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1404 | 1404.1931_arXiv.txt | Most stars form in a clustered environment. Therefore, it is important to assess how this environment influences the evolution of protoplanetary discs around young stars. In turn, this affects their ability to produce planets and ultimately life. We present here for the first time 3D SPH/N-body simulations that include both the hydrodynamical evolution of the discs around their natal stars, as well as the dynamics of the stars themselves. The discs are viscously evolving, accreting mass onto the central star and spreading. We find penetrating encounters to be very destructive for the discs as in previous studies, although the frequency of such encounters is low. We also find, however, that encounter influence the disc radii more strongly than other disc properties such as the disc mass. The disc sizes are set by the competition between viscous spreading and the disruptive effect of encounters. As discs spread, encounters become more and more important. In the regime of rapid spreading encounters simply truncate the discs, stripping the outer portions. In the opposite regime, we find that the effect of many distant encounters is able to limit the disc size. Finally, we predict from our simulations that disc sizes are limited by encounters at stellar densities exceeding $\sim 2-3 \times 10^3 \ \mathrm{pc}^{-2}$. | Stars form in regions of enhanced ambient gas and stellar densities compared to the Galactic field \citep{lada03}. Whether or not these density peaks are long-lived or disperse on a dynamical time (i.e.~whether they become bound stellar clusters or unbound associations) depends crucially on their initial densities and the resulting star formation efficiencies \citep{kruijssen12}. In Milky Way-like galaxies, about $10$\% of all stars are born in bound stellar clusters \citep{bastian08}, but this number increases with the gas surface density to up to $\sim50$\% in high-density starburst environments \citep{goddard10,adamo11,kruijssen12d,silvavilla13}. The cluster environment leaves a spectacular imprint on the star formation process. Through the collective feedback of young stars such as stellar winds and photoionising radiation, natal gas is ejected and the accretion discs surrounding protostars may be destroyed by external photoevaporation \citep{adams04,pelupessy12,dale13}. Combining the current observational and theoretical understanding of planet-, star-, cluster- and galaxy formation, \citet{longmore14} estimate that some 10\% of all stars in the Universe may have the formation of planets (or lack thereof) in their habitable zones affected by their natal cluster environment. In this paper, we concentrate on the dispersal of gas from protoplanetary discs through encounters with neighbouring stars. It serves as a first step to obtaining a detailed understanding of how the cluster environment affects the evolution of protoplanetary discs. In isolation, an effective viscosity causes the redistribution of angular momentum within the gaseous disc \citep{lyndenbellpringle}. This leads to disc spreading on the one hand and mass accretion onto the central star on the other hand. While the latter process is routinely observed \citep[e.g.,][]{Gullbring98,Natta2004,2008ApJ...681..594H,2012ApJ...755..154M}, there are only a few observational reports of disc spreading \citep{Isella2009,2011A&A...529A.105G}, as high spatial resolution is needed to resolve the disc size. Within the current limitations, these works show how simple theoretical models are able to reproduce the observed rate of disc spreading. In practice, these works have concentrated on the nearest star-forming regions, namely Taurus-Auriga and Ophiucus, which are characterized by a lower stellar density when compared with more crowded regions, like the Orion Nebula Cluster (ONC). After several Myr of this slow and quiet evolution, it appears that another destructive process kicks in, and the disc is rapidly cleared on a $\sim 10^5 \ \mathrm{yr}$ timescale \citep{2010ApJS..186..111L,2011MNRAS.410..671E,2013MNRAS.428.3327K}. Currently, internal photoevaporation is the best candidate mechanism for such a fast disc dispersal \citep{UVswitch,AlexanderEvol,2009ApJ...705.1237G,Owen1st,Owen11Models}. Does a clustered environment impact this picture of disc evolution? \citet{2012A&A...546L...1D} found evidence of a dependence of the observed disc sizes on the environmental surface stellar density. In particular, discs in crowded environments, that is, at stellar densities above $10^{3.5}\,\mathrm{pc}^{-2}$, are systematically smaller than their counterparts in less crowded fields. Observationally, it is known that proximity to high mass stars may alter the evolution of protoplanetary discs via external photoevaporation \citep{1998AJ....115..263O,2010ApJ...725..430M,2012ApJ...757...78M}. The high-energy radiation from massive stars can ionize and evaporate the material in the atmosphere of discs even at distances of $\simeq 1 \ \mathrm{pc}$ \citep{Johnstone98,adams04}. Although there are spectacular images of this process in silhouette discs (proplyds) in the Orion Nebula Cluster (ONC), overall this process is not expected to be the main driver of disc evolution \citep{2010ARA&A..48...47A}. Another important process occurring in a clustered environment are stellar encounters. Most of the previous work done on this problem has concentrated on modeling a given existing stellar cluster. These studies \citep{2001MNRAS.325..449S,2006ApJ...642.1140O,2008A&A...487L..45P,2012ApJ...756..123O,2013ApJ...769..150C} involved either semi-analytic solutions or pure N--body simulations in which close stellar encounters are recorded and the effect of single encounters on a putative disc is inferred \emph{a posteriori} (using results from simulations with pure N-body techniques, or including also hydrodynamical effects). \cite{2001MNRAS.325..449S} performed N--body simulations using 4 000 stars in virial equilibrium in an $r^{-2}$ density distribution with a half--mass radius of $\sim1$ pc to model the ONC. The ONC is a popular target, being the nearest massive star--forming region, where many protostellar discs are observed in silhouette against the bright nebula \citep{ricci2008,robberto13}. The ONC contains $\sim 4 000$ stars (i.e. $\sim2\times10^{3}$ M$_{\odot}$) in a $\sim$5 pc diameter volume, has a one--dimensional velocity dispersion of 2.5 km s$^{-1}$ and a core density of 4.7$\times$10$^{4}$ pc$^{-3}$. \cite{2001MNRAS.325..449S} found that $\sim8\%$ of all stars and $\sim30\%$ of core stars suffered a sub--100 $\mathrm{au}$ encounter after 12.5 Myr of integration and concluded that encounters were unlikely to significantly affect the disk population. However, they cautioned that the sharp outer cutoff in their stellar distribution caused their models to expand significantly. This naturally lowers the encounter rate. \cite{2006ApJ...642.1140O} used very similar initial conditions to \cite{2001MNRAS.325..449S} in their N--body study of the ONC, except that they also modelled sub-virial clusters. Instead of recording the single closest encounter, as in \cite{2001MNRAS.325..449S}, \cite{2006ApJ...642.1140O} recorded the complete encounter history of objects on the following grounds: (a) the closest encounter may not be the most destructive, since a distant flyby of a massive perturber can do more damage than a near--miss with a low--mass object (\cite{2007ApJ...656..275M} found that unequal--mass encounters are more destructive); (b) some stars will experience several encounters whose effects may be cumulative. They also concluded that the fraction of stars experiencing sub-$100$ \au encounters in 12.5 Myr was small, at most $\sim12\%$. They estimated disk mass--losses explicitly using a fitting formula from an extension of the parameter--space study of \cite{2005A&A...437..967P} and found that serial encounters and flybys of massive perturbers were able to affect the disk population, at least in the dense core of the cluster. They concluded that, over 12.5 Myr, $\sim4\%$ of disks in the ONC and $\sim10\%$ of disks in the core would be destroyed outright, assuming initial disk radii of 100 \au. This fraction is increased to $\sim9\%$ and $\sim20\%$ respectively if initial outer disc radii of 200 $\mathrm{au} $ are assumed instead. \cite{2008A&A...487L..45P}, again considering the ONC, pointed out that close encounters involving disc--bearing stars in clusters can also result in bursts of accretion due to spiral arms induced in the disks. They concluded that this is a common phenomenon in dense cluster cores, driving accretion rates up by orders of magnitude for short periods ($10^{2}$--$10^{4}$ yr), during which 5--10$\%$ of the disk may be accreted. \cite{2008A&A...492..735P} speculated that such events may be observed as FU Orionis outbursts. \cite{2012ApJ...756..123O} studied star--disk interactions in the Arches cluster. The Arches is more massive ($\sim3\times10^{4}$M$_{\odot}$), more compact (with a half--mass radius of $\sim0.4$ pc) and therefore much denser ($\sim2\times10^{5}$ pc$^{-3}$) than the ONC. It also has a higher one--dimensional velocity dispersion (5.4 km s$^{-1}$). Encounter rates are therefore expected to be substantially higher in this system and, since its age is comparable to that of the ONC (a few Myr), the total number of encounters that have already occurred should also be much higher. Observations by \cite{2010ApJ...718..810S} revealed that the disc fraction in the Arches cluster is an increasing function of distance from the cluster centre, rising from a few percent in the core to around ten percent at a radius of 0.3 pc. \cite{2012ApJ...756..123O} again employed N--body modelling, and post--processing with techniques similar to \cite{2006ApJ...642.1140O} to infer disk mass--losses. They found disc destruction fractions of 10$\%$ in the entire cluster and 30$\%$ in the core over 2.5 Myr. \citet{malmberg2007,malmberg2011} performed N-body simulations of clusters containing a number of stars ranging from 150 to 1000 and half-mass radii ranging from 0.38 to $7.66 \ \mathrm{pc}$. They quantified from the simulations the fraction of singletons, which they define as those stars that never had encounters closer than $1000 \ \mathrm{au}$. They found that in some cases almost $\sim 85 \%$ of stars are non-singletons, with potential impact on planet-forming protoplanetary discs and already existing planetary systems. They also found frequent exchange of stars in binaries. The effect of fly-bys on already formed planetary systems is to lead to planet ejection and eccentricity excitation in planets that are left in the system, as well as increasing the probability of planet-planet scattering after the fly-by. These authors note that due to binary heating, which will lead to a significant cluster expansion, most encounters happen when the cluster is very young, and therefore the impact on proto-planetary discs can be significant. \cite{2013ApJ...769..150C} performed N--body simulations of a set of idealized, fractally--substructured clusters. Since the local stellar density in cluster subgroups can greatly exceed the average density, the encounter rates in a structured cluster should be considerably higher than in a smooth cluster with otherwise comparable properties. However, stellar subgroups are dynamically erased on a crossing time in bound systems, so it is not obvious that the total \emph{number} of encounters will be higher in a structured cluster. \cite{2013ApJ...769..150C} found that the overall enhancement in the number of encounters due to substructure is only a factor of a few, and that discs in such clusters are not likely to be significantly dynamically influenced in this way. In this paper, we present results from hybrid N-body - smoothed particle hydrodynamics (SPH) simulations of coupled cluster and protoplanetary disc evolution. Therefore, we do not need to infer \emph{a posteriori} the effect of encounters on discs, but we compute it self-consistently together with the stellar dynamics. This allows us to include effects that were neglected in previous studies: \begin{itemize} \item disc spreading and truncation by encounters; \item accretion onto the central star; \item the finite time for a disc to regain equilibrium after an encounter; \item the inclination of the rotation axis with respect to the inclination of the two stars' orbital plane, which has an important effect (it is well known, for example, that a retrograde passage is much less harmful for the disc than a prograde one); \item disc-disc interactions, if both stars in an encounter have a disc; \item the mass transfer between stars, possibly leading to the formation of a new disc. \end{itemize} Rather than trying to accurately reproduce one particular stellar cluster, we concentrate here on an idealized model. This allows us to work in a controlled environment, identifying the new phenomena that arise due to the new computational method adopted. At this stage, we are able to make some preliminary comparison with observations. The questions we want to answer are: \begin{itemize} \item How important are stellar encounters for disc dispersal? \item What are the conditions under which disc sizes are set by stellar encounters? \item Are there observables in protoplanetary discs that can tell us if a disc or a disc population experienced significant encounters? \end{itemize} Our paper is organized as follows. After describing the computational method in section \ref{sec_model}, we present our results in section \ref{sec_results}. We discuss them in section \ref{sec_discussion}, comparing with results from a simple semi-analytical method and with observations, and we draw our conclusions in section \ref{sec_conclusions}. | \label{sec_conclusions} We have presented results from the first hybrid N-body - SPH simulations of coupled cluster and protoplanetary disc evolution. The discs in our simulation are expanding and accreting material onto the star due to viscous evolution, but they are also affected by close encounters between stars. Our simulations allow us to study whether a clustered environment, through the effect of encounters, modifies the protoplanetary disc evolution. We find that encounters can be very destructive for some of the discs, leading to almost complete dispersal for some of them. However, overall the median mass of the discs is not severely affected by the encounters. We find that disc size is much more affected by encounters than disc mass. In the case in which disc spreading is fast, due to a high viscosity, only close encounters matter, as any mass redistribution in the disc caused by more distant encounters is quickly washed out. In this case, the close encounters simply truncate the disc at a given radius. If instead the spreading is not fast enough, we find a regime where distant encounters can have a significant impact on the discs, hardening their surface densities, and thus shrinking their radii. This also makes the discs more resistant to mass stripping by subsequent encounters. Therefore, we stress the importance of hydrodynamical numerical simulations of this kind to yield accurate predictions of the impact of stellar encounters on disc sizes. Finally, we confirm that theoretically we expect to see a cut-off at stellar densities higher than $10^{3.5}\,\mathrm{pc}^{-2}$ in the disc sizes due to the effect of encounters. Further work is needed to probe the high stellar densities present in real stellar clusters. | 14 | 4 | 1404.1931 |
1404 | 1404.4725_arXiv.txt | In this paper, we discuss the light-curve features of various flaring scenarios in a time-dependent leptonic model for low-frequency-peaked blazars. The quasar 3C273 is used as an illustrative example. Our code takes into account Fermi-II acceleration and all relevant electron cooling terms, including the external radiation fields generally found to be important in the modeling of the SEDs of FSRQs, as well as synchrotron self absorption and $\gamma\gamma$ pair-production. General parameters are constrained through a fit to the average spectral energy distribution (SED) of the blazar by numerically solving the time-dependent Fokker-Planck equation for the electron evolution in a steady-state situation. We then apply perturbations to several input parameters (magnetic field, particle injection luminosity, acceleration time scale) to simulate flaring events and compute time-dependent SEDs and light curves in representative energy bands (radio, optical, X-rays, $\gamma$-rays). Time lags between different bands are evaluated using a discrete cross correlation analysis. We find that Fermi-II acceleration has a significant effect on the distributions and that flaring events caused by increased acceleration efficiency of the Fermi II process will produce a correlation between the radio, optical and $\gamma$-ray bandpasses, but an anti-correlation between these three bandpasses and the X-ray band, with the X-rays lagging behind the variations in other bands by up to several hours. | \label{intro} Blazars represent a class of radio-loud Active Galactic Nuclei that consists of BL Lac objects and Flat Spectrum Radio Quasars (FSRQs). The spectral energy distributions (SED) of blazars is characterized by two broadband, nonthermal components that span from the radio to UV or X-ray wavelengths and from x-rays to high-energy $\gamma$-rays. The extreme inferred isotropic-equivalent $\gamma$-ray luminosities, combined with rapid variability in different bandpasses, in some cases, down to just a few minutes, provides evidence for strong Doppler boosting in these sources. This is considered to be the result of beamed emission from relativistic jets closely alisgned with our line of sight. It is generally accepted that the low-energy spectral component is synchrotron emission of relativistic electrons/positrons. For the origin of the high-energy SED component, two different approaches have been discussed, referred to as leptonic and hadronic models \citep[for a review of both types of models, see, e.g.,][]{boettcher07,Boettcher12}. In the leptonic scenario, the X-ray to $\gamma$-ray emission is due to the inverse Compton scattering off the relativistic electrons, with the target photon fields either being the synchrotron photons within the emission region (SSC = synchrotron self Compton), or photons external to the jet (EC = external Compton). The external photon fields can include the accretion disk \citep{Dermer92,Dermer93}, the broad line region (BLR), \citep{Sikora94,Blandford95}, or even an infra-red emitting dust torus (IR) that surrounds the central accretion flow onto supermassive black hole \citep{Blazejowski00}. Leptonic models are widely used and have been relatively successful in modeling the SEDs and some variability features of blazars. In hadronic models \citep[e.g.,][]{mb92,Mastichiadis95,Muecke01,Muecke03,Mastichiadis05,Boettcher13}, $\gamma$-rays are the result of proton synchrotron radiation as well as $\pi^0$-decay and synchrotron and Compton radiation from secondary particles in photo-pion induced cascades, presuming the existence of ultrarelativistic protons in the emission region. While such models have also had success in modeling the SEDs of blazars and remain viable, rapid variability observed in blazars is more readily explained in terms of the much shorter acceleration and cooling time scales of relativistic leptons. Therefore, in this work, we will focus on leptonic models.\\ The shapes of the spectral components provide insight into the underlying particle distribution that is producing the emission. Simple power-law and broken power-law electron distributions with parameters chosen ad-hoc, have often been invoked in order to model the SEDs of blazars. Alternatively, log-parabolic electron distributions have been successfully employed to produce the curved synchrotron and Compton spectra observed in many blazars \citep{Massaro04,Massaro06,Cerruti13,Dermer14}. The log-parabola function is characterized by two variables that describe the spectral parameter of the electron distribution and the spectral curvature of the distribution. The log parabolic shape has been shown to be analytically related to a stochastic acceleration mechanism, in which the acceleration probability decreases with energy \citep{Rani11,Massaro06}. Such a connection of log-parabolic spectra and acceleration mechanisms naturally arises in solutions of the time dependent Fokker-Planck equation that contains a momentum diffusion term, indicative of Fermi II acceleration, when the evolution reaches equilibrium \citep{Tramacere11,Massaro06}. They showed that the spectral curvature is inversely proportional to the momentum diffusion coefficient, since the diffusion term acts to broaden the shape of the particle distribution. \\ Second order Fermi acceleration is therefore a viable mechanism for producing log-parabola particle spectra which may be hard enough to reproduce the hard spectra of $\gamma$-ray emission observed in several TeV blazars \citep{Lefa11,Asano13}. It has also been shown that relativistic Maxwellian electron distributions can result from stochastic acceleration processes balanced by radiative losses \citep{1Schlickeiser84}. For the full time dependent Fokker-Planck equation incorporating Fermi II acceleration, general solutions have been found using Green's functions and the application of spectral operators \citep{Stawarz08,Tramacere11}. Solutions to the Fokker-Planck equation incorporating both Fermi I and Fermi II processes, have been developed for the application of the transport of energetic ions \citep{Becker06}. Solutions have also been obtained that consider both Fermi I and Fermi II acceleration, and radiative losses in the Thomson regime \citep{1Schlickeiser84,2Schlickeiser84}. However, when Klein-Nishina effects on the electron cooling rates, as well as absorption processes in the radiation transfer problem, are taken into account, one needs to resort to numerical solutions of the Fokker-Planck equation. \cite{Asano13} developed a time dependent Leptonic model that incorporated Fermi II processes to study the hard spectrum of the blazars Mrk 421 and 1ES 1101-232. The curvature of the electron spectrum, as well as the hard $\gamma$-ray spectra could be reproduced by a model that utilizes a stochastic momentum diffusion process (Fermi II). \\ In this paper, we use a time-dependent Leptonic model that incorporates Fermi acceleration and self-consistent radiative losses, including synchrotron and Compton scattering on internal (SSC) and external (EC) radiation fields as well as synchrotron self-absorption and $\gamma\gamma$ absorption and pair production. The purpose of this paper is to investigate the effects of various flaring scenarios, including Fermi-II acceleration, in external-Compton dominated blazars to complement the study for SSC-dominated sources by \cite{Asano13}. Therefore, while our code is applicable to all types of blazars, we here focus on its application to FSRQs. We describe the model and underlying assumptions in Section \ref{theory}. We use our code to study the influence of Fermi-II acceleration on the quasi-equilibrium particle distribution and light-curve features, including possible time delays beween variations in different frequency bands. These features are studied with parameters motivated by an SED fit to the FSRQ 3C 273, described in Section \ref{spectrum}. Once we have obtained appropriate baseline parameters, we choose a set of input parameters (specifically, the particle injection luminosity, the magnetic field, and the acceleration time scale) to perturb them in the form of a Gaussian in time, in order to study the light curves in the radio, optical, x-ray and $\gamma$-ray bandpasses (Section \ref{lightcurve}). In Section \ref{correlation}, we perform a discrete correlation function analysis on the light curves obtained in the preceding section, to determine possible time lags between the selected bandpasses. We summarize and discuss our results in Section \ref{results}. Throughout this paper, a cosmology with $\Omega_{m} = 0.3$, $\Omega_{\Lambda} = 0.7$, and $H_{0} = 70 km s^{-1} Mpc^{-1}$ is used. | 14 | 4 | 1404.4725 |
|
1404 | 1404.3619_arXiv.txt | Historically, light curve studies of supernovae (SNe) and other transient classes have focused on individual objects with copious and high signal-to-noise observations. In the nascent era of wide field transient searches, objects with detailed observations are decreasing as a fraction of the overall known SN population, and this strategy sacrifices the majority of the information contained in the data about the underlying population of transients. A population level modeling approach, simultaneously fitting all available observations of objects in a transient sub-class of interest, fully mines the data to infer the properties of the population and avoids certain systematic biases. We present a novel hierarchical Bayesian statistical model for population level modeling of transient light curves, and discuss its implementation using an efficient Hamiltonian Monte Carlo technique. As a test case, we apply this model to the Type~IIP SN sample from the \PS\ Medium Deep Survey, consisting of \NphotTot\ photometric observations of \NIIPtotalF\ SNe, corresponding to a joint posterior distribution with \totalparams~parameters under our model. Our hierarchical model fits provide improved constraints on light curve parameters relevant to the physical properties of their progenitor stars relative to modeling individual light curves alone. Moreover, we directly evaluate the probability for occurrence rates of unseen light curve characteristics from the model hyperparameters, addressing observational biases in survey methodology. We view this modeling framework as an unsupervised machine learning technique with the ability to maximize scientific returns from data to be collected by future wide field transient searches like LSST. \smallskip | \label{sec:intro} The majority of luminous transients in the universe are core-collapse supernovae (CC-SNe), marking the explosive deaths of massive stars \citep{Heger03,Smartt09}. Stellar evolution theory, as well as both detailed observations of the explosive transient and fortuitous pre-explosion observations of the progenitor star, point to progenitor initial mass as the primary factor determining stars' eventual death state. Metallicity, rotation rate, binarity, and other properties play important secondary roles, and permutations of these parameters are likely responsible for the extreme diversity of core-collapse supernovae phenomenology observed in the universe \citep{Heger03,Smartt09,Smith11,Ekstrom12,Jerkstrand13}. The progenitor star mass distribution for each SN~type, as well as the distribution of these secondary factors, have far reaching implications throughout astrophysics, influencing the theory of stellar evolution \citep{Groh13}, galactic chemical evolution \citep{Nomoto13}, hydrodynamic feedback in galaxy formation \citep{Stilp13}, and astrobiology \citep{Lineweaver04}. Studies of individual transients typically focus on well observed cases within each object class, capitalizing on the availability of detailed and high signal-to-noise observations to facilitate comparison to finely tuned hydrodynamic explosion simulations and analytic light curve models (e.g. \citealt{Mazzali03,Utrobin08}). Syntheses of these observations, studies of large samples of SNe of a given class, are then typically composed of samples culled from these well observed cases (see e.g. \citealt{Nomoto06,Bersten09,Jerkstrand13}). However, the properties of luminous and/or high signal-to-noise objects within a survey sample may be systematically different from their lower luminosity / signal-to-noise counterparts, and traditional targeted transient searches themselves are inherently biased towards particular SN progenitor properties like high metallicity \citep{SandersIbc,nes2010ay}. To derive truly robust and unbiased inferences about SN progenitor populations, it is therefore necessary to study transient samples in a fashion as complete and observationally agnostic as possible. Here we discuss a methodological framework for the simultaneous modeling of multi-band, multi-object photometric observations from wide field transient surveys, which addresses certain biasing factors inherent to transient searches. This method is rooted in ``hierarchical'' and ``multi-level'' Bayesian analysis, where information about similar events within a sample is partially pooled through a hierarchical structure applied to the joint prior distribution (see \citealt{BDA3} and references therein; see \citealt{Mandel09} for applications to SN~Ia light curves). We adopt Hamiltonian Monte Carlo as a computational technique to efficiently explore the high-dimensional and strongly correlated posterior distribution of this hierarchical model \citep{Betancourt13}. The result of this modeling is simultaneous inference on physically-relevant light curve parameters describing individual objects in the sample, as well as the parameter distribution among the population, regularized by the application of minimal (``weakly informative'') prior information. In Section~\ref{sec:model} we discuss the design and implementation of a hierarchical Bayesian model capable of simultaneously fitting large quantities of raw photometric data from wide field transient surveys to infer the population properties of the underlying SN sample. We test this model with a sample dataset of Type~IIP SNe from the \PS\ (PS1) survey (Section~\ref{sec:data}), previously published in \cite{Sanders14IIP}. We explore the results of this test in Section~\ref{sec:results}, including comparison with inferences drawn from traditional modeling based on fits to individual light curves. We discuss the implications of this methodology for future wide field transient surveys in Section~\ref{sec:disc} and conclude in Section~\ref{sec:conc}. | \label{sec:conc} We have explored the use of Bayesian hierarchical modeling and Hamiltonian Monte Carlo (HMC) to enable population-level inference on multi-band transient light curves from comprehensive analysis of optical photometry from wide field transient searches. The primary conclusions of this work are: \begin{itemize} \item While computational limits still challenge the implementation of hierarchical models, due to the high curvature in their joint posterior distributions, sufficient convergence is achieved in the bottom level model parameters (Section~\ref{sec:res:converge}) to enable their immediate application for transient light curve studies. \item Comparisons between light curve posterior predictive distributions from our hierarchical model fit to the individual light curve fits of \cite{Sanders14IIP} show strong agreement for well identified parameters, and show an advantage for hierarchical models among poorly identified parameters (Section~\ref{sec:res:PPC}). In particular, partial pooling of parameter information between transients supports improved regularization of light curve shapes, and supports model selection between partially degenerate light curve parameter scenarios. \item By directly modeling the underlying transient population, hierarchical models permit inference on the occurrence of properties not observed within the dataset (Section~\ref{sec:res:pop}). This feature is of particular value in overcoming observational biases induced by ground based transient searches, such as the under-representation of long duration transients like some SNe~IIP. \end{itemize} We have concluded with a discussion of future directions for this modeling (Section~\ref{sec:disc}), including applications to upcoming wide field transient searches, extensions to the hierarchical model structure developed here, and expanded capabilities to be enabled by the advent of Riemannian manifold Hamiltonian Monte Carlo. | 14 | 4 | 1404.3619 |
1404 | 1404.3569_arXiv.txt | The \textit{Solar Dynamics Observatory/Helioseismic and Magnetic Imager} (SDO/HMI) filtergrams, taken at six wavelengths around the Fe \textsc{i} 6173.3 \AA \ line, contain information about the line-of-sight velocity over a range of heights in the solar atmosphere. Multi-height velocity inferences from these observations can be exploited to study wave motions and energy transport in the atmosphere. Using realistic convection simulation datasets provided by the \textsf{STAGGER} and \textsf{MURaM} codes, we generate synthetic filtergrams and explore several methods for estimating Dopplergrams. We investigate at which height each synthetic Dopplergram correlates most strongly with the vertical velocity in the model atmospheres. On the basis of the investigation, we propose two Dopplergrams other than the standard HMI-algorithm Dopplergram produced from HMI filtergrams: a line-center Dopplergram and an average-wing Dopplergram. These two Dopplergrams correlate most strongly with vertical velocities at the heights of 30\,--\,40 km above (line-center) and 30\,--\,40 km below (average-wing) the effective height of the HMI-algorithm Dopplergram. Therefore, we can obtain velocity information from two layers separated by about a half of a scale height in the atmosphere, at best. The phase shifts between these multi-height Dopplergrams from observational data as well as those from the simulated data are also consistent with the height-difference estimates in the frequency range above the photospheric acoustic cutoff frequency. | In recent helioseismology studies multi-height velocity and intensity information have been used in addition to standard photospheric Dopplergrams. \inlinecite{2008A&A...481L...1M} investigated the phase shift between photospheric- and chromospheric-intensity datasets obtained by the \textit{Hinode} satellite. They reported large phase differences along the $p$-mode ridges and no phase difference on the $f$-mode ridge. \inlinecite{2009ApJ...694L.115N} inferred chromospheric downflows by interpreting multi-height observations. More recently, \inlinecite{2012SoPh..281..533H} examined the phase differences between several observables originating from various layers obtained by the \textit{Helioseismic and Magnetic Imager} (HMI: \opencite{2012SoPh..275..207S}) and \textit{the Atmospheric Imaging Assembly} (AIA: \opencite{2012SoPh..275...17L}) onboard the \textit{Solar Dynamics Observatory} (SDO: \opencite{2012SoPh..275....3P}). \inlinecite{2012SoPh..tmp..303R} exploited multi-height HMI and AIA data to study power enhancement around active regions at various heights in the atmosphere. They used not only the standard HMI-algorithm Dopplergrams but also Doppler information derived from the line wing (this is similar to what we define as the ``far-wing" Dopplergram in Section \ref{sec:DopDef}). Multi-height information is, however, useful not only for helioseismology studies but also for many other research purposes, for example, the study of energy transport in the solar atmosphere (\textit{e.g.} \opencite{2006ApJ...648L.151J}, \opencite{2008ApJ...681L.125S}, \opencite{2009ASPC..415...95S}, \opencite{2010ApJ...723L.134B}, \opencite{2011A&A...532A.111K}). If we could obtain multi-height velocity information from SDO/HMI datasets, we would have huge datasets available compared with other current instruments; HMI obtains full-disk datasets without significant interruptions. Here we show that we can obtain multi-height velocity information from SDO/HMI observations; we use realistic numerical convection simulations to characterize these multi-height Dopplergrams. In Section \ref{sec:makeDop} we introduce HMI observables as well as the simulation datasets, and define several types of Dopplergrams. In Section \ref{sec:syn_cor} we investigate the contribution heights of the Doppler velocities using realistic convection simulations. On the basis of these contribution height estimates and availability of the observables we choose a pair of Doppler velocities as rather robust multi-height velocity datasets; this is summarized in Section \ref{sec:multi-height}. We measure the phase difference derived from the HMI observation datasets as well as simulated datasets in Section \ref{sec:phase}. Finally, a brief summary is given in Section \ref{sec:Conclusions}. | \label{sec:Conclusions} We confirm that we can obtain multi-height line-of-sight velocity information in the solar atmosphere from SDO/HMI filtergrams. We suggest average-wing and line-center Dopplergrams as well as the general pipeline product, (first) HMI-algorithm Dopplergrams as rather robust multi-height velocity datasets among the several Doppler velocities defined in Section \ref{sec:makeDop} based on the estimate of the contribution layer heights and the availability of the observables. In general, multi-height Doppler observations have the potential to help constrain the height variations of the $p$-mode eigenfunctions in the solar atmosphere (see \inlinecite{2012ApJ...760L...1B}, for applications). In addition, multi-height Doppler observations may be helpful for distinguishing convective motions from oscillations, which in turn may be useful to improve the signal-to-noise ratio in helioseismology studies. We estimate the contribution layer heights of several synthetic Doppler velocities computed using numerical convection simulation datasets, \textsf{STAGGER} and \textsf{MURaM}. Although the contribution layer is rather broad, the contribution layer heights of the average-wing and the line-center Dopplergrams are 30\,--\,40 km lower and 30\,--\,40 km higher compared to the standard HMI-algorithm Dopplergrams, respectively. Note that the height difference between these Dopplergrams is 70 km at most, which is about half, at best, of the scale height around the surface (see Figure \ref{fig:wac}). We can obtain multi-height information from these observables, but since we use the filtergrams taken around a single absorption line, the height difference is relatively limited. HMI observations show clear phase differences between these Dopplergrams at frequencies above the acoustic cutoff frequency. The height difference estimated by the response functions is consistent with the one estimated by the phase differences. HMI observation data also show a clear signature of atmospheric gravity waves in the lower-frequency ranges, while \textsf{STAGGER} simulation data have only a weak signature. Although in this study we limited ourselves to quiet-Sun data for the sake of simplicity, multi-height velocity information in magnetic regions is also of great interest. Since the spectral-line shape is changed not only by the velocity fields but also by magnetic field, it is not straightforward to analyze such observations and one would need radiative-transfer calculations including the effect of magnetic field. The formation layer of the SDO/AIA 1600 \AA \ and 1700 \AA \ passbands are estimated in the lower chromosphere around 430 and 360 km above the surface (see \opencite{2005ApJ...625..556F}). Multi-height Doppler observations from SDO/HMI either alone or together with SDO/AIA or \textit{Interface Region Imaging Spectrograph} (IRIS) observations will potentially be also useful to understand how much wave energy is transported in the atmosphere and corona (\textit{e.g.} the review by \opencite{2010LRSP....7....5R}). \begin{acks}[Acknowledgments] BL acknowledges support from IMPRS Solar System School. BL computed the line profiles from the \textsf{STAGGER} data cubes and the response functions using the \textsf{SPINOR} code. We thank Michiel van Noort, Thomas Straus, and Jesper Schou for helpful discussions and comments. KN and LG acknowledge support from EU FP7 Collaborative Project ``Exploitation of Space Data for Innovative Helio-and Asteroseismology" (SPACEINN). LG acknowledges support from DFG SFB 963 ``Astrophysical Flow Instabilities and Turbulence" (Project A1). The HMI data used are courtesy of NASA/SDO and the HMI science team. This work was carried out using the data from the SDO HMI/AIA Joint Science Operations Center Data Record Management System and Storage Unit Management System (JSOC DRMS/SUMS). The NSO/Kitt Peak FTS data used here were produced by NSF/NSO. RS acknowledges support by NASA grant NNX12AH49G and NSF grant AGS1141921. The \textsf{STAGGER} calculations were performed on the Pleiades cluster of the NASA Advanced Supercomputing Division at Ames Research Center. The German Data Center for SDO (GDC-SDO), funded by the German Aerospace Center (DLR), provided the IT infrastructure to process the data. \end{acks} \appendix | 14 | 4 | 1404.3569 |
1404 | 1404.3995.txt | The Earth is known to be depleted in volatile lithophile elements in a fashion that defies easy explanation. We resolve this anomaly with a model that combines the porosity of collisionally grown dust grains in protoplanetary disks with heating from FU Orionis events that dramatically raise protoplanetary disk temperatures. The heating from an FU Orionis event alters the aerodynamical properties of the dust while evaporating the volatiles. This causes the dust to settle, abandoning those volatiles. The success of this model in explaining the elemental composition of the Earth is a strong argument in favor of highly porous collisionally grown dust grains in protoplanetary disks outside our Solar System. Further, it demonstrates how thermal (or condensation based) alterations of dust porosity, and hence aerodynamics, can be a strong factor in planet formation, leading to the onset of rapid gravitational instabilities in the dust disk and the subsequent collapse that forms planetesimals. | Of all meteorite classes, the elemental abundances of CI chondrites most closely match the Solar photosphere, and those meteorites are believed to comprise the most primitive solid material found in our Solar System \citep{Lodders03}. For that reason, these meteorites are held as the standard for the elemental and chemical composition of solids that condensed out of the Solar nebula, before thermal, chemical, or physical processing took their toll. However, the bulk silicate Earth has a significantly smaller mass-fraction of volatile lithophile elements (low condensation temperature and trapped in the mantle, hence measurable) than do CI chondrites: the relative abundance of lithophile elements with condensation temperatures below $1400$\op K decreases with decreasing condensation temperature \citep{Palme00,McDonough03}. That abundance trend is inverted in Solar abundances which show enhancement in those same volatiles relative to solar twins \citep{Mel09}. Assuming that the Earth formed out of solids which at first had Solar composition, this depletion of the volatile elements has been a long outstanding problem in the formation of the Earth, especially since \cite{Humayun95} showed that this decrease in volatiles does not come with an isotopic signature in potassium. Existing theories for this volatile depletion are unsatisfactory, and explaining it is the goal of this paper. Our model combines many disparate aspects of astrophysics and planetary sciences, from FU Orionis style outbursts and dust dynamics, to isotope ratios and the Goldschmidt classification. In light of the breadth of reader backgrounds, we have attempted to provide adequate introductions to the different moving parts before using them in detail. Accordingly, we include an extended introduction to the problem in \Sec{Prob-overview}, detailing the constraints that any theory for the Earth's volatile depletion must match. We provide a narrative description of our model in \Sec{model_overview} to provide context for the introduction to the various sub-processes found in \Sec{particulars}. \Sec{sec_actual} finally combines the individual pieces into a coherent whole, and \Sec{extension} extends the model to other bodies in the Solar System. We conclude in \Sec{conclusions}. | \label{conclusions} We have presented a model for the depletion of Earth's volatiles as measured with respect to Solar abundances. The model relies on the spatial segregation of volatiles in the gas phase with dust grains through the vertical settling of the dust after an FU Orionis type outburst heats the disk adequately to evaporate the volatiles. The heating causes highly porous grains to contract and settle, abandoning much of the volatiles in the upper reaches of the disk, but only after isotopic reequilibration with those volatiles. Finally, the Streaming Instability gathers the settled grains, triggering gravitational instability and leading to direct collapse to planetesimals. Our model explains not only the Earth's volatile depletion, but also the trickier lack of an isotopic signature in the potassium depletions because it provides long enough timescales for the different isotopes to equilibrate independently. This model predicts that collisional agglomeration naturally creates high porosity fluffy dust grains, and that the planetesimals that became Earth formed early %(at the time of the last FU Orionis type event in the proto-Solar nebula) and in situ. The model predicts volatile depletions for all planets formed within the critical annulus inside of which the background disk was cool enough to allow the formation of fluffy dust grains, but FU Orionis type events were hot enough to melt or sinter the dust grains, causing them to become compact and to settle. In the case of the proto-Solar nebula, this extends from about Venus' orbit to just about Mars', although the implications for Mars are weaker because the heating is unlikely to completely contract the dust grains. Our models ties together a large number of processes which still have large systematic uncertainties. We have observed only a handful of FUors, none to completion, and the growth of dust grains of unknown geometry and surface chemistry has been modeled with only gross simplifications. Indeed, it is only recently that the question of icy surfaces has been treated in some detail alongside SiO$_2$ grains. Accordingly, we have described the simplest version of our model to explain it qualitatively, and to show that it holds quantitative promise. Some extensions are unlikely to pose significant problems: planets must aggregate materials from a significant annulus, rather than merely a single position. This will add together multiple depletion patterns, but if a model predicts a depletion pattern that matches the Earth's at one radial position, feeding from an annulus of modest width near that position should not change the final result dramatically. Like all studies of planet formation, our model would be significantly improved by a better understanding of the outcome of dust-dust collisions. Along with studies of chondrites, we would additionally benefit from systematic study of the diffusion of volatile elements through silicates under hot nebular conditions. Finally, our model makes the study of long (month or year) timescale sintering of Solar Nebula minerals important. The experiences of \citet{Poppe03} shows that such studies need to be performed in zero-G because gravity acts to compress materials. We have also assumed that $R=1$\op AU lies outside of the FUor engine, which allows us to neglect the mass flow onto the star that powers the FUor, keeping our material from moving radially on the timescales we consider. The extent of FUor engines is not known (observations have not yet resolved them), but they may well extend beyond $R=1$\op AU \citep{Zhu10}. While our model can adjust for this by moving radially outwards, this will inevitably result in lower temperatures unless the FUor is brighter. However, the luminosity jump we associate with the FUor is on the strong end of observations, so a large outwards change in radial positioning is probably not possible. Finally, planet migration raises strong questions about the link between the initial and final position of a planet which are beyond the scope of this work. Interestingly though, the Earth's depletion pattern suggests that it must have been made reasonably near $1$\op AU. Too much farther out, and even an FUor wouldn't be able to raise the temperature enough to evaporate volatiles, while too close in, and even quiescent disks will be too hot. The implication that collisionally grown dust in protoplanetary disks has high porosity means that thermal reduction of dust grain porosity will be a significant player in planet formation. By growing dust grains in a fractal, high porosity manner, even high mass dust grains remain tightly coupled to the gas, experiencing relatively low collisional velocities: low porosity dust grains of the same mass would experience destructive collisions. Subsequent heating and grain contraction then led to highly settled dust grains that are subject to the streaming instability, leading to direct gravitational collapse of the dust disk to form planetesimals. While we appeal to FU Orionis type events in this model, other sources of heating could play a role. It is difficult to make specific predictions for our model in the Solar System because it is likely that many objects formed during quiescent phases. The large number of meteorite classes suggests that some probably did form close in time to an accretion event, so preserving pre-solar grains during FUors is a problem, albeit one which applies independently of our model. Further, while we believe FUor events likely promote volatile-depleted rocky planet formation in extra-solar, often heavily embedded, systems, observationally testing that hypothesis is not currently possible in part because of the rarity of FUors. Nonetheless, our model states that an initial supply of highly porous dust grains will be compactified. This should have observational implications, particularly when applied to icy grains on large enough orbits that their their effect on disk spectra occur in annuli which can be resolved. This will require an adequate survey of protostellar systems that future FUors can be compared with their previous, quiescent state. | 14 | 4 | 1404.3995 |
1404 | 1404.2273_arXiv.txt | The dark sector of the Universe need not be completely separable into distinct dark matter and dark energy components. We consider a model of early dark energy in which the dark energy mimics a dark matter component in both evolution and perturbations at early times. Barotropic aether dark energy scales as a fixed fraction, possibly greater than one, of the dark matter density and has vanishing sound speed at early times before undergoing a transition. This gives signatures not only in cosmic expansion but in sound speed and inhomogeneities, and in number of effective neutrino species. Model parameters describe the timing, sharpness of the transition, and the relative abundance at early times. Upon comparison with current data, we find viable regimes in which the dark energy behaves like dark matter at early times: for transitions well before recombination the dark energy to dark matter fraction can equal or exceed unity, while for transitions near recombination the ratio can only be a few percent. After the transition, dark energy goes its separate way, ultimately driving cosmic acceleration and approaching a cosmological constant in this scenario. | The present energy budget of the Universe contains two significant components beyond the Standard Model of particle physics -- dark matter and dark energy -- with little definitely known about their nature. Today, they act very differently, with dark matter clustering into galaxies and enhancing gravitational attraction while dark energy appears spread nearly homogeneously throughout space with a tension that causes the acceleration of the cosmic expansion. Here we consider whether they might have been more closely related in the past. Such a concept might be realized if dark matter and dark energy arose from the same, or related high energy physics processes. Indeed, such connections might arise from decaying moduli in string theory, e.g.\ see \cite{1401.4364} for such a case connecting dark matter and dark radiation. Modified gravity can also cause evolution of the couplings to and between different sectors, e.g.\ the early transition model of \cite{1312.3361}. In particular, the dark sector could have a different character at high redshift, and dark energy could have contributed dynamically at early times, perhaps with density at certain epochs comparable to that of dark matter. Several high energy physics origins for dark energy, such as Dirac-Born-Infeld scalar fields \cite{dbi1,dbi2,dbi3} or dilatons \cite{dilaton1,dilaton2}, predict such early dark energy, and in many cases it acts in a decelerating manner, possibly scaling as the dominant component of energy density, or simply like dark matter. Moreover, such models often involve a non-relativistic sound speed of perturbations. Thus, such cold, early dark energy can act substantially like cold dark matter. For probing the early universe, measurements of the cosmic microwave background (CMB) offer the best evidence, and have already been used to place percent-level limits on dark energy at recombination \cite{planck16,hls}. Here we investigate whether viable models exist in which the early dark energy density at some prerecombination epoch can be of order (or even greater) than the dark matter density, while possessing many of the same properties. Since the dark energy should today be accelerating and fairly smooth, this requires an evolution in its behavior in both equation of state and sound speed. Moreover, so as not to disagree with formation of galaxies and clusters, the dark energy must quickly fade away from the early universe into the matter dominated era where structure grows. Recently, \cite{13052209} investigated a model where such a transition occurred after recombination. However, they kept the fluctuation sound speed in the dark energy to be the speed of light, reducing the effect of perturbations, and adopted a purely phenomenological model for the density evolution. Because the model here behaves like dark matter in both the expansion and perturbations at times before the transition, then if the transition occurred after recombination such additional energy density would just look like added dark matter and be faced with the usual CMB constraints on the dark matter density. Therefore we concentrate on the more interesting case for our model of a prerecombination transition and adopt for the dark energy the barotropic aether \cite{scher04,linscher}, a model with useful and interesting properties. In Sec.~\ref{sec:baro} we explore the physical effects of the barotropic aether as it evolves from dark matter-like behavior at early times to cosmic acceleration at late times. We confront the model with recent CMB data in Sec.~\ref{sec:cmb}, and discuss what this may teach us about the dark sector in Sec.~\ref{sec:concl}. | \label{sec:concl} The dark sector of the Universe presents us with multiple, fundamental mysteries. Observations concentrate at the present epoch, where two quite distinct components appear: clustering, pressureless dark matter and highly smooth, strongly negative pressure dark energy. Within the visible sector of the Universe, the Standard Model of particle physics teaches us that apparently distinct entities can be unified at high energies, corresponding to early times in cosmic history. We have explored a phenomenological model for such a merging of dark matter and dark energy at early times, where barotropic aether dark energy has the properties of dark matter, perhaps through some unspecified direct interaction. At late times this behavior is gone, releasing the two components to evolve very differently. A barotropic aether model has the desired properties of naturally appearing like dark matter ($w=0=c_s^2$) at early times, and then a very rapid evolution away, toward a late time attractor with $w=-1$ and $c_s^2=1$, acting like a cosmological constant. This would give added, rich structure to the dark sector and, if confirmed, a substantial clue to high energy physics. We show that the transition in this model cannot take place arbitrarily rapidly, but must take longer than a number of e-folds $\tau\ge1/3$. Confronting this model with current data, we find that such a model is wholly acceptable if the transition occurs sufficiently before recombination. For example, the asymptotic early ratio of dark energy to dark matter density can be larger than unity if the transition is at $a_t<10^{-3.84}$, or larger than $100$ for $a_t<10^{-4.5}$. Later transitions however are severely constrained by data, especially the CMB temperature power spectrum. We find that the early dark energy to dark matter density ratio cannot exceed 5\%, similar to other early dark energy models, for transitions much after recombination. Even later transitions have been constrained by other work, e.g.\ \cite{13052209}. Even so, a ratio $R=0.025$ at, say, $a_t=10^{-3}$ can contribute energy density interpreted in terms of an effective number of extra neutrino species of $\neff=0.25$. Thus early dark energy remains of interest. An interesting generalization is to consider a whole spectrum of barotropic fields, with all allowed values of sound speed $c_s^2=[0,1]$ (see Appendix~\ref{apx:spec}). This scenario has intriguing properties, with each component dominating in sequence and then fading away, similar to isotopes with different half lives. Because of the physical constraint $c_s^2\ge0$ the late time universe is left with only the $c_s^2=0$ component of dark energy, corresponding to the barotropic aether considered here. However, while somewhat attractive as a way to avoid naturally a coincidence that dark energy only dominates today, it does suffer from increased fine tuning (unless a way can be found to cancel the positive and negative contributions). If the dark sector does come together at high energies, we might expect the transition epoch not to be at eV scales ($\log a_t\approx -3$), but at GeV or higher scales. As this is above the primordial nucleosynthesis scale, observational constraints are lacking. Future work will explore whether inflation -- a very early dark energy period -- can constrain or benefit from such ``mimic'' dark energy. | 14 | 4 | 1404.2273 |
1404 | 1404.2790_arXiv.txt | CCD light curves of the Algol type eclipsing binaries DP Cep, AL Gem, FG Gem, UU Leo, CF Tau and AW Vul were analysed using the Wilson-Deninney code and new geometric and absolute parameters were derived. Due to cyclic apparent orbital period changes of the systems, probably caused by the Light-Time Effect, the contribution of a third light was taken into account in the light curve solution. All the reliable timings of minima found in the literature were used to study the period variations and search for the presence of a tertiary component in the systems. A comparison between the parameters of the third body derived from the light curve and orbital period analyses is also discussed. Moreover, the absolute parameters of the eclipsing binary components were also used to determine their current evolutionary state. | \label{1} The main purpose of this work is the derivation of the light curve (hereafter LC) and orbital period parameters of each system, and a description of the components' current evolutionary stage. The systems were selected due to their \emph{\textbf{O}bserved -- \textbf{C}alculated} times of minima periodic variations (hereafter O -- C). The LC observations were performed in order: (a) to investigate the photometric presence of a tertiary component near the eclipsing binary (hereafter EB) via a third light contribution, (b) to make an approximate calculation of the absolute parameters of the components and (c) to search for possible pulsational behaviour in the systems which are candidates for including a $\delta$ Sct component. According to the O -- C solution, it is feasible to detect which physical mechanisms play a role to the period modulation, such as a third body or the mass transfer between the components or mass loss from the system or magnetic activity. On the other hand, from the LC solution it is possible to determine the Roche geometry of the EB, such as semi-detached, detached or contact configuration. A plausible conclusion is that these solutions are qualitatively connected, even though the methods used for the derivation of their results are different. The LC solution is based on the model of the photometric variation of the EB, while the O -- C solution counts on the variability of its orbital period. Obviously, the LC is an instant image of the EB's long life, while the period changes correspond to a larger time scale. However, although the period variability duration of the system is small (order of decades) comparing with its total lifetime, one is able to unify the information provided from both analyses and obtain a consistent picture of the EB. The existence of a third body orbiting an eclipsing pair is not always easy to be proved. The tertiary component has to be of the same order of magnitude as the EB, in order to be detected photometrically through its light contribution to the total luminosity of the triple system. However, if the third body is an evolved star (e.g white dwarf) it is obvious that it will affect the EB's period, but not the LC. On the other hand, one can also imagine a different situation when only a weakly bound tertiary contributes a non-negligible amount of the third light, but no period modulation is detectable even on longer time scales. Concluding, in the present study we selected the following six cases showing interesting variation of minima times in the O -- C diagram. The results derived from LC and O -- C analysis for each system and their probable connection will be discussed. \subsection{Individual systems} \textbf{DP Cep:} This system (V=12.9 mag, P$\sim$1.27 days) was discovered by \citet{b51} who determined its period. \citet{b52} obtained the first photographic LC of the EB, reported a more accurate period and, according to its B-V index, they listed its spectral type as F0. \citet{b08} classified the system as a semidetached with mass ratio of 0.6, and classified its components as F0 and G0 for primary and secondary, respectively. The first CCD LC was obtained by INTEGRAL (INTErnational Gamma-Ray Astrophysics Laboratory) mission \citep{b31} using the OMC (Optical Monitoring Camera), and \citet{b53} performed the first LC analysis. \textbf{AL Gem:} The light variability of this system (V=9.77 mag, P$\sim$1.39 days) was reported by \citet{b01} and the first photoelectric LC was obtained by \citet{b02}. The system was classified as F4-F6 type star by \citet{b04}, while \citet{b03} measured its colour indexes. \citet{b05}, \citet{b06}, \citet{b22} and \citet{b21} calculated the absolute parameters of the system's components. The spectroscopic survey of \citet{b07} classified the EB as a F5 type, while \citet{b08} referred its primary and secondary components as F5 and K7 type stars, respectively. The first CCD LC was observed by the ASAS project \citep{b17}, but it is poorly covered. Finally, AL Gem is concerned as a candidate triple system \citep{b09}, but its hypothetical tertiary component has not been detected so far. \textbf{FG Gem:} \citet{b10} discovered the variable nature of this EB (B=12.6 mag, P$\sim$0.82 days). Except for the observations for minima timings determination, no complete LC of the system existed until early the decade of 2000. The first LC was obtained by the ASAS project \citep{b17}, while \citet{b11} obtained B and R filter LCs and presented the first geometrical solution. The first approximate spectral classification, namely F7 for primary and K3 for secondary components, respectively, and the derivation of their absolute parameters were made by \citet{b21}. The spectral type of FG Gem is still uncertain since many catalogues refer different colour indexes estimates. In particular, \emph{Lick NPM2 Catalogue} \citep{b12}: B -- V=0.25, \emph{All-sky Compiled Catalogue of 2.5 million stars} \citep{b16}: B -- V=0.835 and \emph{Tycho-2 Catalogue} \citep{b14}: B -- V=0.02. In the catalogue of \emph{T$_{eff}$ and metallicities for Tycho-2 stars} \citep{b15}, except the B -- V indexes, the V -- K index and the metallicity were also taken into account for the classification. According to this catalogue, although the errors are large, the most trustable temperature estimation of the system as T$\sim$7700 +4800/-590 K was provided, and, in addition, the distance of the system as 223 +283/-154 pc was estimated. \textbf{UU Leo:} The eclipsing status of this system (B=12.1 mag, P$\sim$1.68 days) was mentioned by \citet{b18} and its period was given by \citet{b19}. Many timings of minima exist so far, but a complete LC was unavailable. \citet{b22} and \citet{b20} referred the system to be of A2 spectral type. The results of \citet{b21} showed that this result corresponds to the primary component, and in addition they found that the EB's secondary is of G1 spectral class. UU Leo is referred as candidate for triplicity \citep{b09} and also as candidate system for including a pulsating component \citep{b23}. \citet{b24} obtained an O -- C analysis of the EB, and concluded that the mass transfer between the components and magnetic phenomena are the responsible mechanisms for its orbital period changes. \textbf{CF Tau:} Light variations of this EB (V=10.05 mag, P$\sim$2.76 days) were firstly announced by \citet{b25}, while \citet{b26} obtained the first visual LC. \citet{b27} published the first accurate ephemeris of the system, \citet{b03} derived its Str\"{o}mgen colour indexes and \citet{b22} calculated its absolute parameters based on the photometric parallaxes. \citet{b03} classified the system as G0 type, while the same spectral type is also referred by the \emph{Henry Draper Catalogue identifications for Tycho-2 stars} \citep{b29} and \emph{T$_{eff}$ and metallicities for Tycho-2 stars} \citep{b15} catalogue, where, in addition a distance of 159 +170/-79 pc was calculated. In agreement with this spectral values, \citet{b21} estimated the absolute parameters of the EB and classified its components as G0 and K2, for primary and secondary, respectively. \citet{b28} performed photoelectric UBV observations of the EB and calculated its colour indexes in various phases. LITE was proposed as the most plausible explanation for the period changes of the system according to the O -- C analysis of \citet{b30}. A partial CCD LC of the EB by the ASAS \citep{b17} and a complete one by SWASP \citep{b32} projects were obtained. \textbf{AW Vul:} The discovery of the system (V=11.1 mag, P$\sim$0.81 days) was made by \citet{b33}. \citet{b19} and \citet{b27} reported its first ephemerides. \citet{b22}, \citet{b20} and \citet{b21} calculated the absolute parameters of the system, and they are in agreement for the spectral type of the primary component as F0. The visual LC of the EB was obtained by \citet{b34}, while the only existed CCD LC is poorly covered by \citet{b35}. \citet{b36} classified the system also as F0 type and determined its distance as 440 pc. In addition, in the catalogue of \emph{T$_{eff}$ and metallicities for Tycho-2 stars} \citep{b15} the distance of the system is given as 155 +367/-277 pc. Finally, AW Vul is listed in the catalogue of possible triples \citep{b09} and its hotter component is candidate for pulsations \citep{b23}. | \label{7} Complete LCs of the eclipsing systems AL Gem, UU Leo and AW Vul were obtained, and together with the ones of DP Cep, FG Gem and CF Tau obtained from sky surveys were analysed in order to derive their geometrical and absolute parameters. In addition, the LC residuals of all systems, except for CF Tau, were tested for possible pulsations but the results were negative. O -- C analysis of each system was performed and showed probable existence of a tertiary component orbiting the common center of mass. In all cases, the contribution of a third light to the total luminosity was taken into account in the LC analysis, and its fraction $L_{3,LC}(\%)$ (see Table \ref{tab2}) was derived. According to the minimal mass of the third body found from the O -- C analysis, one is able to derive its absolute luminosity by assuming its MS nature and by using the mass-luminosity relation for such stars (L$\sim$M$^{3.5}$); then, the calculation of its light contribution to the total luminosity of the system, given the absolute luminosities of the eclipsing components $L_1$ and $L_2$ (see Table \ref{tab3}), is feasible. The following formula gives the expected luminosity fraction of the third companion: \begin{equation} L_{3, O-C}(\%)=\frac{100\cdot M_{3,min}^{3.5}}{L_1+L_2+M_{3,min}^{3.5}} \end{equation} Alternatively, since a cool component of an EB is potentially able to present magnetic activity, one may consider the variation of the quadrupole moment $\Delta$Q and examine its possible influence on the observed period changes as suggested by \citet{b48}. According to the equations of \citet{b49} and \citet{b50}, respectively: \begin{eqnarray} \Delta P &=& A \sqrt{2 (1 - \cos \{2\pi P/P_{3}\})},\\ &&\nonumber \\ \frac{\Delta P}{P} &=& -9 \frac{\Delta Q}{Ma^2} \,\,\,\ \end{eqnarray} and the calculated absolute parameters of the possible magnetically active component (see Table \ref{tab3}), one is able to test this mechanism also as an alternative explanation for the cyclic period variations. According to the criterion suggested by \citet{b50}, the cyclic period changes of an EB could be caused by magnetic influences of a component if its quadrupole moment lies between the range $10^{50}<\Delta Q<10^{51}$. The eclipsing pair DP Cep is described by a semidetached configuration, where the cooler and more evolved component fills its Roche Lobe. Its primary is located at the upper limit of the MS, indicating that its Hydrogen-burning era has almost come to the end. Although we expected, except the LITE curve, a parabola imposition due to the possible mass transfer, the O -- C analysis did not confirmed it. The long gap of observations between 1952-1975, or the slow mass transfer rate seem to be the best explanations for this result. A third light was detected in the LC analysis ($\sim$5\%), while the minimal mass of the hypothetical MS third companion suggests a light contribution of $\sim$2.5\%. This small difference can be easily overcome, by suggesting an inclination of about 65$^\circ$ for the third body's orbit, which yields a mass of $\sim$0.95 $M_\odot$ capable to produce the observed additional light. Another possible explanation for the cyclic period modulation could be the magnetic activity of the secondary component. Its quadrupole moment was found to have a value of 6.6$\times$10$^{50}~g\cdot~cm^2$, therefore its magnetic influences are potentially candidate formers of the period changes. However, the existence of this mechanism was not detected in the LC analysis (e.g. O'Connell effect), thus other methods of verification (e.g. more LCs, spectroscopy, polarimetry or long term variations of minima timings studies) are needed. AL Gem is in a detached configuration with the hotter component at the edge of MS and the cooler one more evolved. The LC analysis yielded a third light contribution of $\sim$12.5\%, while the expected one for a MS star with minimal mass $\sim$0.28 $M_\odot$ is about 0.19\%. This discrepancy might be explained by assuming the non-coplanar orbits of the tertiary component and the EB, similarly to DP Cep. Moreover, the angular separation between the third body and the EB was found to have a relatively large value (98 mas), while the magnitude difference turned out as 2.2 mag. On the other hand, a cool photospheric spot was revealed through the LC analysis, but the application of Applegate's mechanism yielded for the secondary component of the system a value of $\Delta$Q$<10^{50}~g\cdot~cm^2$, therefore is not able to explain the period changes. Spectroscopic and probably also interferometric observations, given the high brightness of the system, will certainly solve the mystery of the existence of the third component. The system FG Gem was found in semidetached status with the secondary component filling its Roche Lobe and consisting of one MS component and an evolved one. Unlike the previous case, the observed and the expected light contribution from a tertiary component are in perfect agreement (L$_3\sim$4.8\%), hence the system can be considered as a triple, with its components in almost coplanar orbits. However, the O -- C analysis did not reveal any sign of parabolic behaviour (mass transfer-loss indicator), so very probably the mass transfer has not started yet, or its rate is very slow to be detected in a $\sim$80 $yrs$ time span. The value of quadrupole moment was found for the cool component as 2.5$\times$10$^{50}~g\cdot~cm^2$, and can also cause the observed cyclic changes, but this hypothesis need further observations to verify the existence of magnetic activity. UU Leo is a detached EB, with the hotter component close to the ZAMS limit indicating its young age and the cooler one beyond the TAMS edge, probably in the subgiant era. Excellent agreement was found between L$_{3,LC}$ and L$_{3,O-C}$ resulted in $\sim$11.3\%. However, a $\Delta$Q value of 6.9$\times$10$^{50}g\cdot~cm^2$ was derived for the secondary component, and, in addition, a cool spot presence was also needed for better fit in the LC analysis. Both mechanisms describe well the O -- C data. As in the case of AL Gem, the relatively high value of L$_3$ is very promising to be detected by spectroscopy which is expected to give definite answer for the mechanism forming the orbital period of the eclipsing pair. The geometrical status of CF Tau was found to be a detached one. Its components are located beyond the TAMS line, while according to their almost equal masses and positions in the M -- R diagram, they seem to follow similar evolutionary tracks. The observed third light contribution yielded as 4.3\%, while a nearby MS star having a mass value of $\sim$0.89 $M_\odot$, as it was found from the O -- C analysis, is capable to produce this additional luminosity (L$_{3,O-C}$=6.5\%). Thus, since these values are both qualitatively and marginally quantitatively in agreement, the triplicity of the system is established. Both components are potential candidates for magnetic activity, but neither the LC revealed such behaviour nor the value of quadrupole moment ($\Delta$Q$<10^{50}~gr\cdot~cm^2$) is able to explain the cyclic period modulation. AW Vul can be considered as a classical Algol, since its cooler, more evolved and less massive component fills its Roche Lobe, while the hotter one is a typical MS star. The mass transfer mechanism was not detected in the O -- C analysis, but likely to FG Gem it is possible that its rate is very slow to play a significant role in such a relatively small observational time coverage ($\sim$80 $yrs$). No evidence of magnetic activity was found in the LC analysis, and in addition the Applegate's mechanism is insufficient to describe the cyclic variations of the orbital period, since a value of $\Delta$Q$<10^{50}~g\cdot~cm^2$ was yielded for the cooler component. On the other hand, the observed additional light was found as L$_{3,LC}\sim$3\%, while the expected one from a minimal mass of 0.24 $M_\odot$ was L$_{3, O-C}\sim$0.1\%. This difference can be explained either by assuming the non MS nature of the third body and by inserting a giant age hypothesis or that its inclination is $\sim$25$^\circ$ instead of coplanar orbit, which, according to mass function, yields a value of M$_{3}\sim$0.7 $M_\odot$. Such a mass value satisfies the observed third light in the case of a MS star. Radial velocity measurements are needed for deriving the absolute parameters of these close binaries more conclusively, while in the cases of AL Gem and UU Leo, where a significant third light was traced, the spectroscopic detection of the third body will verify the current results. New times of minima for all systems will help to achieve a better coverage of the cyclic changes of their orbital periods and will probably reveal new variations which are not so obvious in the current O -- C data sets. Moreover, a prospective interferometric detection of the third body would be also very profitable, with the system AL Gem being the most suitable one for such observations. | 14 | 4 | 1404.2790 |
1404 | 1404.0276_arXiv.txt | We study the dynamical properties and space distribution of dark energy in the weak and strong gravitational fields caused by inhomogeneities of matter in the static world of galaxies and clusters. We show that the dark energy in the weak gravitational fields of matter density perturbations can condense or dilute, but amplitudes of its perturbations remain very small on all scales. We illustrate also how the ``accretion'' of the phantom dark energy onto the matter overdensity forms the dark energy underdensity. We analyze the behavior of dark energy in the gravitational fields of stars and black holes with the Schwarzschild metric. It is shown that, in the case of stars, the static solution of the differential equations for energy-momentum conservation exists and describes the distribution of density of dark energy inside and outside of a star. We have found that for stars and galaxies its value differs slightly from the average and is a bit higher for the quintessential scalar field as dark energy and a bit lower for the phantom one. The difference grows with the decrease of the effective sound speed of dark energy and is large in the neighborhood of neutron stars. We obtain and analyze also the solutions of equations that describe the stationary accretion of the dark energy as a test component onto the Schwarzschild black hole. It is shown that the rate of change of mass of the dark energy is positive in the case of quintessential dark energy and is negative in the case of the phantom one. | Astrophysicists have enough evidence that our Universe is expanding with positive acceleration, instead of deceleration, as it should be in a world filled only with ordinary matter ruled by usual laws of gravity. The reality of this fact is beyond doubt, although its interpretation is far from having a single meaning (see, for example, books and review articles \cite{GRG2008,Amendola2010,Wolchin2010,Ruiz2010,Novosyadlyj2013m} devoted to this problem). One of the most developed interpretations in the terms of theoretical modeling, comparing of numerical predictions with observational data, and determination of the parameters and their confidence intervals is a scalar field that fills the Universe almost homogeneously and slowly rolls down to the minimum of its own potential in the case of a quintessential scalar field or slowly rolls up to the maximum in the case of a phantom one. There are many possible realizations of such a field (one can find a large portion of references in the sources mentioned above), which is why the additional observational or experimental tests, which could help to constrain the number of candidates at least in the class of scalar field models of dark energy, are needed. It seems that observational cosmological data obtained in the last few years \cite{wmap9a,Planck2013b,Percival2010,Anderson2012,Padmanabhan2012,wigglez,6dF,snls3,union,Riess2009,Steigman2007,Wright2007} prefer phantom dark energy \cite{Novosyadlyj2012,Novosyadlyj2013,Xia2013,Cheng2013,Rest2013,Shafer2013,Novosyadlyj2014,Hu2014}. However, we are still far from establishing the physical nature of dark energy, and it is necessary to search for new sensitive tests for it. The properties of dark energy are extensively investigated in the FRW space-time as the cosmological background and a lot of models have been proposed. The physical parameters of dynamical dark energy (energy density, potential etc.) are time dependent in such a space-time. And contrarily, the properties of dynamical dark energy in the static world of gravitationally bound systems are unknown. We try to remove this lack by analysis of dynamical properties of the scalar field as the dark energy of either quintessential or phantom type in the galaxies and vicinity of stars and black holes. The goal of this paper is to study the distribution of the density of quintessential and phantom dark energy in the weak and strong gravitational fields of the static astrophysical objects. In this regard we are interested Refs. \cite{Babichev2004}, the authors of which drew the conclusion about decreasing the mass of a black hole onto which the phantom dark energy falls. To verify this conclusion, we analyze the evolution of density perturbations of dark energy in the weak and strong gravitational fields, the possibility of static configurations of the fields in neighborhood of massive astrophysical objects, and the accretion of the dark energy onto them. We prove the illusiveness of decreasing the black hole mass caused by accretion of the phantom dark energy. | The analysis of dynamics of dark energy in the static field of a gravitationally bound systems has shown that density of dark energy in the galaxy clusters, in the galaxies, and in the vicinity of and inside the stars is by few orders smaller than the mean density of dark matter. It has shown that quintessential ($-1<w<-1/3$) and phantom ($w<-1$) dark energy in the static world of galaxies and clusters of galaxies is gravitationally stable: under the influence of self-gravitation, it can only oscillate. In the gravitational fields of dark matter perturbations, it can inflow monotonically, but the amplitudes of perturbations of density and velocity remain small at all scales interesting for astrophysics. It has also been shown that the accretion of phantom dark energy in the region of overdensities of dark matter causes the formation of the negative perturbation of the density of dark energy. The behaviour of dark energy in the gravitational field of static nonrotating spherical objects has been studied. The static distribution of dark energy density in their vicinity and inside them has been obtained [Eqs. (\ref{rho_stat2})-(\ref{rho_stat7})]. It has been found that the dark energy density differs very slightly from the mean one; the magnitude of relative deviation of the density, $\delta_{de}=\rho_{de}(0)/\overline{\rho}_{de}-1$, has maximum in the center and depends on the EoS parameter $w_{\infty}$, and the effective sound speed $c_s$ of dark energy as well as on the gravitational radius of central body $r_g$ and its relative size $\alpha^{-1}=R/r_g$. Its sign is positive for the quintessential scalar field and negative for the phantom one. For stars similar to the Sun, $\delta_{de}\sim10^{-7}$, and for neutron stars, $\sim10^{-1}$. When $c_s\sim1$ and $w\sim-1\pm0.1$, the density inside and near the compact object is small and barely differs from the average value. The difference becomes significant when $c_s\rightarrow 0$ or $r\rightarrow r_g$. The solutions of equations describing the stationary accretion of dark energy as a test component onto the Schwarzschild black hole have been obtained [Eqs. (\ref{rho_ac1})-(\ref{ac_rho_0.5})] and analyzed. It was shown that the mass of dark energy, which crosses the sphere of radius $r>r_g$ toward the center, is determined by the parameters of dark energy and the square of the black hole event horizon. The rate of change of mass is a positive quantity in the case of quintessential dark energy and negative in the case of the phantom one. | 14 | 4 | 1404.0276 |
1404 | 1404.5723_arXiv.txt | {WZ Sagittae is the prototype object of a subclass of dwarf novae, with rare and long (super)outbursts, in which a white dwarf primary accretes matter from a low mass companion. High-energy observations offer the possibility of a better understanding of the disk-accretion mechanism in WZ Sge-like binaries.} {{ We used archival {\it XMM}-Newton and Swift data to characterize the X-ray spectral and temporal properties of WZ Sge in quiescence.}} {{ We performed a detailed timing analysis of the simultaneous X-ray and UV light curves obtained with the EPIC and OM instruments on board {\it XMM}-Newton in 2003. We employed several techniques in this study, including a correlation study between the two curves. We also performed an X-ray spectral analysis using the EPIC data, as well as Swift/XRT data obtained in 2011.}} {We find that the X-ray intensity is clearly modulated at a period of $\simeq 28.96$ s, confirming previously published preliminary results. We find that the X-ray spectral shape of WZ Sge remains practically unchanged between the {\it XMM}-Newton and Swift observations. However, after correcting for inter-stellar absorption, the intrinsic luminosity is estimated to be ${\rm L^{Una}_X=(2.65\pm0.06)\times 10^{30}}$ erg s${\rm ^{-1}}$ and ${\rm L^{Una}_X=(1.57\pm0.03)\times 10^{30}}$ erg s${\rm ^{-1}}$ in 2003 and 2011, respectively. During the Swift/XRT observation, the observed flux is a factor $\simeq 2$ lower than that observed by {\it XMM}-Newton, but is similar to the quiescent levels observed various times before the 2001 outburst.} {} } | A cataclysmic variable (CV) is a binary system {consisting of a white dwarf primary} which accretes matter from a low mass companion via Roche lobe overflow (for a review, {see} \citealt{warner1995}). Systems with {a primary with} a relatively low magnetic field ($\ut<$0.1 MG) are {expected to accrete} via a Keplerian disk. In {such a} case, half of the total potential gravitational energy is dissipated by the viscosity, with the {remainder being} radiated away by the boundary layer. {The} spectral energy distribution emitted by the accretion disk peaks in the optical and ultraviolet bands, while the boundary layer radiates predominantly in the extreme ultraviolet and X-rays. Typical X-ray luminosities of CVs are in the range $10^{30}$--$10^{32}$ erg s$^{-1}$ (see e.g. \citealt{lamb82}, \citealt{baskill}, \citealt{erik}). {\it XMM}-Newton (\citealt{jansen2001}) is particularly {useful} for studying quiescent CVs as its large effective area allows to detect faint sources {in general and during} dips and eclipses {in particular}. Moreover, the possibility to observe the source simultaneously in the optical {or ultraviolet (UV)} bands with the optical monitor (OM) opens the possibility to study the correlations between light curves of the same source in different wavelengths {taken at exactly the same time}. WZ Sagittae (hereafter WZ Sge) is currently known to be the closest CV ($43.5\pm0.3$ pc, see \citealt{harrison2004}). It {reaches} $V\simeq 7-8$ {during outbursts} (e.g. \citealt{patterson2002}, \citealt{kuulkers2011}); it spends most of the time, however, in a quiescent state characterized by rather modest optical magnitudes in the range $14-16$ (e.g, \citealt{steeghs}, \citealt{kuulkers2011}). It has a {short orbital period of $\simeq 81.6$ min} (Krzemi\'nski 1962; Warner 1976). Apart from showing {a large outburst amplitude}, WZ Sge has also a long outburst recurrence time: it goes into outburst every 20-30 years. {In the literature, there are reports of large outbursts of WZ Sge in 1913, 1946, 1978 and 2001 (see e.g., \citealt{mayall1946}, \citealt{brosh1979}, \citealt{mattei2001}, \citealt{godon2004}, \citealt{ishioka2001} and references therein). For a the historical record of these observations, we refer to \citet{kuulkers2011}. Several observational campaigns were devoted to the study of the source characteristics in detail during the 2001 outburst (see, e.g., } \citealt{patterson2002}, \citealt{knigge2002}, \citealt{long2003}, \citealt{sion2003}). {One prominent scenario for the long outburst recurrence time is that the inner part of the accretion disk is truncated by the magnetic field of the white dwarf (see, e.g., \citealt{warner1995}, \citealt{hameury1997}). In this scenario, the $\simeq 28$ s periodic modulation in the optical data (see, for example, \citealt{patterson1998}, \citealt{lasota1999}), is interpreted as possibly related to the white dwarf spin period. In Sect. 4, however, we discuss a counterargument, suggesting that other scenarios should be considered.} {The component masses of WZ Sge are still not well known. The photometric solution of \citet{smak1993} gives a mass of the white dwarf of $M_{wd}\simeq 0.45$ M$_{\odot}$ and a mass ratio $q\simeq 0.13$, whereas \citet{spruit1998}, who modeled the hot spot at which the mass stream transferred from the companion hits the outer accretion disk, obtain $M_{wd} \simeq 1.2$ M$_{\odot}$ and a mass ratio $q\simeq 0.075$. As shown by phase-resolved spectroscopy (\citealt{steeghs}), the binary system is characterized by a primary white dwarf with mass in the range $0.88$ M$_{\odot}$--$1.53$ M$_{\odot}$ and a low mass companion of $0.078$ M$_{\odot}$--$0.13$ M$_{\odot}$ which is close to the brown dwarf mass threshold. If the mean velocity of absorption lines is interpreted as being due to gravitational red-shift (and one uses the mass-radius relation), then the mass of the primary {is inferred} to be $(0.85\pm 0.04)$ M$_{\odot}$. In the present work, we use the latter value for the mass of the white dwarf in WZ Sge, i.e. $0.85$ M$_{\odot}$.} WZ Sge has been intensively observed in the X-ray band. \citet{patterson1998} described both the ROSAT and ASCA observations (as well as Einstein and EXOSAT ones) {obtained in quiescence} . This analysis was successively re-done by \citet{gun2005} who reported a quiescent 0.1--2.4 keV flux (as obtained from ROSAT PSPC in 1991) of $\simeq 2.8 \times 10^{-12}$ erg cm$^{-2}$ s${\rm ^{-1}}$ (corresponding to a luminosity of $\simeq 6.3\times 10^{29}$ erg s${\rm ^{-1}}$ for a distance of $\simeq 43.5$ pc). In addition, \citet{hasenkopf}, re-analyzing a 1996 ASCA observation of WZ Sge, found a 0.5--10 keV flux of $\simeq 4.7\times 10^{-12}$ erg cm$^{-2}$ s${\rm ^{-1}}$, thus implying a luminosity of $\simeq 1.0\times 10^{30}$ erg s${\rm ^{-1}}$. Furthermore, the 2001 outburst of WZ Sge was observed in X-rays (see, {e.g.,} \citealt{wheatley2001}, \citealt{kuulkers2002}, \citealt{wheatley2005}). In this paper we present the result of $\sim 9.9$ ks {\it XMM}-Newton and $\sim 1.4$ ks Swift observations of WZ Sge acquired in 2003 and 2011, respectively, i.e. almost two and ten years after the {most recent} outburst. {The {\it XMM}-Newton data were already reported by \citet{mukai2004}, who used a multi-temperature plasma to describe the observed WZ Sge X-ray spectra and found a 2-10 keV band flux of $\simeq 7.0\times 10^{-12}$ erg cm${\rm ^{-2}}$ s${\rm ^{-1}}$ (i.e., much larger than the flux inferred by using the 1996 ASCA data). We, here report on a coherent periodicity of $\simeq 28.96$ s in the same {\it XMM}-Newton observation, when using all the information down to $0.2$ keV. The detected periodicity is close to that found in the optical reported by \citealt{mukai2004}.} The paper is structured as follows: in Sect. \ref{s:xmm1} and \ref{s:swift1} we present the available data and give details about the {\it XMM}-Newton and Swift data reduction, respectively; in Sect. \ref{s:result} (and related sub-sections) we present the results of our timing and spectral analysis. Finally, in Sect. \ref{s:conclusion} we conclude on our observations. | \label{s:conclusion} \subsection{Quiescent X-rays} In this paper, we presented the analysis of an archival {\it XMM}-Newton observation in 2003 (for a preliminar study see \citealt{mukai2004}) and newly acquired Swift data in 2011 of WZ Sge. {WZ Sge's X-ray spectral properties in the 0.2-10 keV energy band remained practically unchanged between the 2003 and 2011 observations. We estimated an unabsorbed intrinsic luminosity of ${\rm L^{Una}_X=(2.65\pm0.06)\times 10^{30}}$ erg s${\rm ^{-1}}$ and ${\rm L^{Una}_X=(1.57\pm0.03)\times 10^{30}}$ erg s${\rm ^{-1}}$ for the 2003 and 2011 observations, respectively. The luminosity in 2011 is a factor $\simeq 2$ lower than that in 2003, indicating that WZ Sge returned to a level similar to that observed prior to the last source outburst in 2001. The high-energy light curves confirm the existence of a dip close to the orbital phase $\simeq 0.7$. Although this feature is not strong, it also appears in the softness ratio light curve, similar to that seen using ROSAT/PSPC data (\citealt{patterson1998}). Dip structures in the light curves are naturally explained in the framework depicted by \citet{frankkinglasota1987} (see also \citet{smak1971}) which is supported by the numerical simulation of \citet{hirose1991} and \citet{armitage1998}. The model found its application in explaining periodic orbital dip features in the high-energy light curves of nova-like systems (see e.g. \citealt{hoard2010} and \citealt{nucita2011}) and of magnetic white dwarfs (see \citealt{ramsay2009}). According to this model, once the mass flow reaches the inferior conjunction at the orbital phase 0.7, part of the accreting matter sets sufficiently high above the disk, thus obscuring the white dwarf and producing the observed dip. \subsection{Periodicities at 27.87 and 28.96 s} With an improved analysis (using all the available data down to $0.2$ keV) we find a coherent periodicity of $\simeq 28.96$ s in the 2003 observation. This confirms the weak detection reported by \citet{mukai2004}; the period is close to that found in optical data reported by the same authors. We did not detect the $27.87$ s oscillation attributed to the white dwarf spin (see e.g. \citealt{patterson1998}, \citealt{lasota1999}), similar to \citet{mukai2004}. } The origin of the 27.87 s {\sl and\/} the 28.96 s periods in WZ Sge has been a long-standing puzzle. For example, \citealt{robinson1978} interpreted these two, distinct, periods as due to non-radial pulsations. \citet{patterson1998} cautiously argued that the 27.87 s period seen in the ASCA X-ray data was the spin period of a magnetic white dwarf. \citet{welsh2003} presented a balanced review of the two models, and pointed out the difficulties with both. Given that 10 years have elapsed since then, during which a large body of recent observations of non-radial pulsations in other low-accretion rate dwarf novae have been obtained, we present a re-assessment of the models. \citet{lasota1999} proposed that the 27.87 s is the white dwarf spin period, while the 28.96 s signal is due to reprocessing of the spin signal by a blob at the outer rim of the Keplerian disk. While this explanation is viable for an {\em optical} modulation at the 28.96 s period, it fails to explain the 28.96 s {\em X-ray} period. While intermediate polars often show X-ray spin and sideband signals simultaneously (\citealt{norton}), this is believed to be due to stream overflow -- mass transfer stream that skirts the surface of the disk and is directly captured by the magnetic field of the white dwarf. It is hard to see how the white dwarf can accrete directly from a blob at the {\sl outer\/} edge of the disk. If, instead, the Keplerian period at the inner edge of the disk is 733.5 s (see, however, objections to this idea by \citealt{lasota1999}), this could in principle lead to an X-ray modulation at the 28.96 speriod. In this case, however, it would be difficult to avoid a strong X-ray modulation at the 733.5 s period (\citealt{wynn1992}), given the high inclination of the WZ Sge system. That is, when the blob that feeds the magnetic pole is on the Earth side of the white dwarf, the pole that is facing the Earth would accrete more favourably. This is likely to lead to a higher observed X-ray flux than when the blob is on the far side. In addition, both the inner and outer radii of an accretion disk are not constant when accretion rate varies; it is not clear how a blob with a Keplerian period of 733.5 s is always favoured, when both the optical (\citealt{kuulkers2011}) and the X-ray (this work) brightness show secular variability. Further arguments against an intermediate polar model come from the hardness curves. We do not find any evidence for X-ray spectral variations along the 28.96 s signal. Usually, in intermediate polars the $X$-ray emission is softer when brighter. We note that, indeed, WZ Sge is not classified as standard member of this class of objects (see e.g. \citealt{knigge2002}). The above described weaknesses, however, may not be fatal for the magnetic CV model of the twin periods. Nevertheless, the $XMM$-Newton detection of the 28.96 s signal makes the argument that the 27.87 s period is the spin period of a magnetic white dwarf somewhat weaker. There is little doubt that the short period variability seen in another faint CV, GW Lib (see also below), is due to non-radial g-mode pulsations of the white dwarf that dominates its optical light in quiescence (\citealt{vanzyl2004}). Since then, similar pulsations have been discovered in about a dozen of other faint, white dwarf-dominated CVs (see, e.g., \citealt{szkody2010} and references therein). In the case of these accreting white dwarfs which are rapidly rotating and have peculiar abundances, these pulsations are more complicated than in the non-accreting ZZ Cet stars. For example, CV primaries may show pulsations outside the ZZ Ceti instability strip. The $<$30 s periods in WZ Sge, however, are significantly shorter than those seen in GW Lib type CVs ($>$200 s). Moreover, X-rays are generated by accretion, and it is not clear how non-radial pulsations would modulate the X-ray flux. In summary, the origin of the twin pulsations is as mysterious as ever. The long-term stability of the intermittent 27.87 s period remains a strongest argument for this to be the spin period of the white dwarf, but this leaves us without a clear understanding of the 28.96\,s period. It is interesting to note, that \citet{Mukadam2013} found several puzzling features in the pulsational variability of another CV, EQ Lyn. In addition to possible ways to reconcile these observations with our understanding of g-mode pulsations, they considered alternatives models: r-mode pulsations and accretion disk pulsations. We should maybe keep in mind such alternative possibilities when considering WZ Sge. \subsection{On the quiescent rate of accretion} In addition to the twin periods of 27.87 s and 28.97 s, WZ Sge possesses several characteristics that made it stand out among dwarf novae. These include the short orbital period, the quiescent spectrum which is dominated by the white dwarf photosphere, the large outburst amplitude and the long inter-outburst interval. However, recent advances show that CVs with many of these latter characteristics are in fact quite common. In particular, the Sloan survey has revealed a large population of CVs near the period minimum (P$<$ 88 min) whose spectra are often dominated by the white dwarf photosphere (\citealt{Gaensicke}). The {earlier} surveys did not go deep enough to show the prevalence of this population. Many of these newly discovered systems are candidate WZ Sge stars in terms of their outburst characteristics -- they are generally seen in a quiescent dwarf nova-like state since their discovery, so any outbursts must be infrequent. The best studied such system is the aforementioned CV, GW Lib, whose discovery in fact predated the Sloan survey. Its well-documented 2007 outburst (\citealt{Byckling}; \citealt{Vican}) is the second known after the discovery outburst in 1983. It has a 76.8 {min} orbital period, its quiescent spectrum is dominated by the white dwarf photosphere, the outburst amplitude is large ($\sim$9 mag), and its duration long ($\sim$26 day). Surely, GW Lib presents a similar challenge to the disk instability model that WZ Sge does. Yet, despite intensive observations motivated by its status as the prototype CV with non-radial pulsations, no spin-period signature has ever been observed in GW Lib. Of the many systems that share various degrees of similarity with WZ Sge (\citealt{Gaensicke}), only V455 And (HS 2331+3905; \citealt{Araujo}) is known to be magnetic. Intensive searches for additional non-radial pulsators have not led to discoveries of magnetic CV signatures among other WZ Sge-like systems. Unless all systems near the period minimum are sufficiently magnetic to create a hole in the disc, yet somehow manage to hide any spin signatures, we must seek an explanation for the long interval, long duration and large amplitude outbursts that do not rely on the primary's magnetic field. In particular, if the correlation found by \citet{patterson2011} between the outburst recurrence time and the mass ratio is confirmed, some factor directly related to the mass ratio is strongly implicated as the cause of the long recurrence time in WZ Sge type systems; the magnetic field of the white dwarf would be a second parameter, not the primary. If that is the case, the twin periods of WZ Sge, whatever their origin, may well be red herring in terms of understanding the outburst properties of WZ Sge. For example, while the detailed propeller model of \citet{matthews2007} can still explain WZ Sge, it fails to explain GW Lib whose outburst properties are similar to those of WZ Sge. On a possibly related note, while the X-ray luminosity of WZ Sge is low compared to dwarf novae with frequent outbursts (U Gem and SU UMa types; \citealt{Byckling2010}), it is higher than that of GW Lib or the Sloan-selected systems studied by \citet{reis2013}. Given this, future studies should strive to understand why the quiescent accretion rate in WZ Sge is high compared to other WZ Sge systems, not why it is lower than in normal dwarf novae. | 14 | 4 | 1404.5723 |
1404 | 1404.3956.txt | The large-scale structure of the Universe formed from initially small perturbations in the cosmic density field, leading to galaxy clusters with up to $10^{15}$~M$_\odot$ at the present day. Here, we review the formation of structures in the Universe, considering the first primordial galaxies and the most massive galaxy clusters as extreme cases of structure formation where fundamental processes such as gravity, turbulence, cooling and feedback are particularly relevant. The first non-linear objects in the Universe formed in dark matter halos with $10^5-10^8$~M$_\odot$ at redshifts $10-30$, leading to the first stars and massive black holes. At later stages, larger scales became non-linear, leading to the formation of galaxy clusters, the most massive objects in the Universe. We describe here their formation via gravitational processes, including the self-similar scaling relations, as well as the observed deviations from such self-similarity and the related non-gravitational physics (cooling, stellar feedback, AGN). While on intermediate cluster scales the self-similar model is in good agreement with the observations, deviations from such self-similarity are apparent in the core regions, where numerical simulations do not reproduce the current observational results. The latter indicates that the interaction of different feedback processes may not be correctly accounted for in current simulations. Both in the most massive clusters of galaxies as well as during the formation of the first objects in the Universe, turbulent structures and shock waves appear to be common, suggesting them to be ubiquitous in the non-linear regime. | \label{intro} %---------------------------------------------------------% %The current model The current hierarchical paradigm of structure formation is set within the spatially flat {\it $\Lambda$-Cold Dark Matter} model \citep[$\Lambda CDM$;][]{Blumenthal_1984} with cosmological constant, also known as the {\it concordance} model. Tight constraints on the parameters of the underlying cosmological model have now been placed thanks to the combination of different observational probes \citep[see, e.g.][for recent reviews]{Voit_2005, Allen2011, Hamilton_2013}. In the resulting scenario \citep[see][and references therein]{Planck_2013_parameters} the Universe, whose age is estimated to be $\sim 13.8$~Gyr, is composed of dark energy ($\Omega_\Lambda\approx0.7$), dark matter ($\Omega_{{DM}}\approx0.25$) and baryonic matter ($\Omega_{b}\approx0.05$), with a Hubble constant given by $H_0\approx67$~km/s/Mpc. In addition, the primordial matter power spectrum seems to be characterized by a power-law index $n\approx0.96$ with an amplitude $\sigma_8\approx0.83$. % Outline of the hierarchical paradigm Within this paradigm, the formation of the first structures in the Universe is driven by the gravitational collapse of small inflation--induced matter density perturbations existing in the primordial matter density field. Predictions from N-body simulations \citep[e.g.][]{Klypin_1983} have confirmed that the growth of these perturbations gives rise to the formation of a complex network of cosmic structures interconnected along walls and filaments concerning a wide range of scales. % First structures in the Universe The first structures in the Universe are expected to form at redshifts of $10-30$ in dark matter (DM) halos of $10^5-10^8$~M$_\odot$ \citep{Tegmark97, Barkana01, Glover05, Bromm09}. A crucial condition for these DM halos to form stars or galaxies is the ability of their gas to cool in a Hubble time. To address this question, \citet{Tegmark97} have modeled the cooling in DM halos of different virial temperatures, showing that a virial temperature of at least $1000$~K is required so that efficient cooling via molecular hydrogen can occur. Such a temperature corresponds to a mass scale of \begin{equation} M_{H_2}\sim10^{6.5}\left( \frac{10}{1+z} \right)^{3/2}M_\odot. \end{equation} Halos of this or slightly higher masses are typically referred to as the so-called minihalos, which are generally assumed to harbor the first primordial stars in the Universe. Their formation has been explored through detailed numerical simulations starting from cosmological initial conditions, following the formation of the first minihalos and their gravitational collapse, including gas chemistry and cooling, down to AU-scales or below \citep{Abel02, Bromm04, Yoshida08}. The first such simulations typically followed only the formation of the first peak during the gravitational collapse, hinting at the formation of rather massive isolated stars of $\sim100-300$~M$_\odot$ due to the rather high accretion rates of $\sim10^{-3}$~M$_\odot$~yr$^{-1}$. % Subsequent studies have explored the formation of self-gravitating disks and their fragmentation at later stages \citep{Stacy10, Clark11, Greif11, Greif12, Latif13a}, indicating the formation of star clusters and binaries rather than isolated stars. The resulting initial mass function (IMF) of these stars is expected to be top-heavy, with characteristic masses in the range of $10-100$~M$_\odot$. The studies involving sink particles further suggest that low-mass protostars can be ejected from the center of the halo via 3-body interactions, thus implying the potential presence of primordial stars with less than a solar mass that could survive until the present day. Radiative feedback seems to imply an upper mass limit of $50-100$~M$_\odot$ \citep{Hosokawa11, Susa13}. In DM halos with virial temperatures above $10^4$~K, an additional cooling channel is present via atomic hydrogen. In such DM halos, also referred to as atomic cooling halos, cooling is always possible via atomic hydrogen lines, helium lines or recombination cooling, while the minihalos may not be able to cool if their molecular hydrogen content is destroyed by photodissociating backgrounds \citep{Machacek01, Johnson07, Johnson08, Schleicher10, Latif11}. Such halos are also more robust with respect to the first supernova explosions \citep{Wise08_3, Greif10}, and may thus give rise to a self-regulated mode of star formation. In the presence of a strong radiative background, for instance from a nearby galaxy, they may remain metal-free and collapse close to isothermally at $\sim8000$~K \citep{Omukai01, Spaans06, Schleicher10, Shang10, Latif11, Latif13c, Prieto13}. While the initial studies followed on the collapse of the first peak \citep{Wise08_2, Regan09a, Shang10}, \citet{Regan09b} aimed at following the longer-term evolution confirming the formation of a self-gravitating disk. \citet{Latif13b} recently pursued the first high-resolution investigation on the fragmentation of such halos on AU scales, finding that fragmentation may occur, but does not inhibit the growth of the resulting central objects. For the high accretion rates of $\sim1$~M$_\odot$~yr$^{-1}$ measured in their simulations, radiative feedback is expected to be negligible \citep{Hosokawa12} and the formation of very massive objects with up to $10^5$~M$_\odot$ seems feasible \citep{Schleicher13}. Such supermassive stars are expected to collapse via the post-Newtonian instability and form the progenitors of supermassive black holes \citep[SMBHs;][]{Shapiro86}. Depending on previous metal enrichment and the ambient radiation field, atomic cooling halos may also gather the proper conditions for the formation of the first galaxies. The formation and evolution of these galaxies, directly connected to the formation of the first stars and their associated radiative or supernova feedback, represent a crucial and complicated aspect of the whole cosmic history. In this sense, the main focus of this review will be on the formation of the first stars and SMBHs, and the reader is referred to the reviews by \citet{Bromm09} and \citet{Bromm11} concerning the formation and properties of the first galaxies. %Galaxy clusters: main properties In the hierarchical paradigm of structure formation, the first objects are the building blocks of subsequent structure formation, leading to larger galaxies and galaxy clusters through accretion and mergers \citep[e.g.][]{Somerville12}. As a consequence of this connection, regardless of the wide range of involved scales, a number of physical processes, such as the generation of turbulence during collapse and the relevance of cooling and feedback processes, seem to be common in the formation of the different cosmic structures. Roughly speaking, the cosmic hierarchy is delimited, in terms of mass and formation time, by the first galaxies in the early Universe and the most massive galaxy clusters at the present day, whereas the bulk of galaxies generally lie in-between these extreme cases. However, a full understanding of galaxy evolution represents a complex and fundamental topic in cosmology that is being currently investigated by a considerable number of authors \citep[see][for a recent review on the current status of galaxy formation]{Silk_2012}. % Given the complexity of this topic and the limited space available for this review, we avoid any description of galaxy evolution. Instead, since we are mostly interested in the role that the physics of plasma plays on the formation of cosmic structures, we will focus both on the formation of the first objects, i.e. the first stars and massive black holes, as well as on the large galaxy clusters at the present day. These extreme scenarios will allow us to illustrate the importance of cooling, turbulence and feedback during structure formation independently of the considered scales. Galaxy clusters are the largest nonlinear objects in the Universe today and thus a central part of the large--scale structure (LSS). Clusters of galaxies, whose total masses vary from $10^{13}$ up to $10^{15}M_\odot$, are characterized by very deep gravitational potential wells containing a large number of galaxies ($\sim 10^2-10^3$) over a region of a few Mpc \citep[see, e.g.][for an early review on galaxy clusters]{Sarazin_1988}. Although most of the mass in clusters is in the form of DM, a very hot and diffuse plasma, the intra--cluster medium (ICM), resides within the space between galaxies in clusters. The ICM, where the thermal plasma coexists with magnetic fields and relativistic particles, holds the major part of the baryonic matter in clusters. This cluster environment affects the evolution of the hosted galaxies by means of a number of dynamical processes such as harassment, ram--pressure stripping or galaxy mergers \citep[e.g. see][for a textbook on galaxy formation and evolution]{Mo_2010}. The intra--cluster plasma, with typical temperatures of $T\sim 10^{7}-10^{8}$ K, strongly emits X-ray radiation, causing clusters of galaxies to have high X-ray luminosities, $L_X\sim 10^{43}-10^{45}$~erg/s. In addition, the ICM is quite tenuous, with electron number densities of $n_e \sim 10^{-4}-10^{-2}$ $cm^{-3}$ and, although it is formed mainly of hydrogen and helium, it also holds a mean abundance of heavier elements of about $\sim1/3$ of the solar abundance. %Role of clusters in Cosmology Given their typical extensions and their deep gravitational potential wells, clusters of galaxies are fundamental for our comprehension of the Universe, marking the transition between cosmological and galactic scales. Whereas on cosmological scales the growth of perturbations is mainly driven by the effects of gravity on the DM component, on galactic scales gravity operates in connection with a number of gas dynamical and astrophysical phenomena. Given such an scenario, galaxy clusters and, in particular, the hot intra--cluster plasma represent a fascinating and complex environment harboring a wide range of astrophysical and dynamical processes related to both the gravitational collapse and the baryonic physics: gravitational shock waves, gas radiative cooling, star formation (SF), gas accretion onto SMBHs hosted by massive cluster galaxies, feedback from supernovae (SNe) or active galactic nuclei (AGN), shock acceleration, magnetohydrodynamical (MHD) processes, gas turbulence, ram--pressure stripping of galaxies, thermal conduction processes, energetics associated to the populations of cosmic ray (CR) electrons and protons, etc. All these processes are manifested by a number of cluster observables such as the thermal X-ray emission, the Sunyaev-Zel'dovich effect \citep[SZ;][]{SZ_1972}, the spectra of galaxies, or the radio synchrotron and gamma-ray emissions associated to the population of non-thermal particles. As a consequence, galaxy clusters reside in an incomparable position within astrophysics and cosmology: while the number and distribution of clusters can be used to place constraints on the current model of cosmic structure formation, a thorough understanding of the complicated processes determining the properties of the hot intra--cluster plasma seems to be crucial to fully understand galaxy cluster observations. %This review... In this review, we describe the formation of the large-scale structure of the Universe in the framework of the $\Lambda$CDM model. A particular focus is both on the formation of the first objects, i.e. the first stars and massive black holes, as well as on the large galaxy clusters at the present day. In both applications, we emphasize the role of gravitational as well as non-gravitational plasma physics such as turbulence, cooling, magnetic fields or feedback processes. The overall structure of this review is as follows: in \S\ref{sec:theory_sf} we start by reviewing the basic concepts of cosmic structure formation, from the early linear evolution of small density perturbations out to the complex collapse of real overdensities; in \S\ref{sec:mf} we overview the relevance for cosmology of a proper calibration of the halo mass function; in \S\ref{sec:early_universe} we describe the formation of the first halos in the early Universe; a brief description of the self-similar model of the intra--cluster plasma is done in \S\ref{sec:self_similar_evolution}, whereas in \S\ref{sec:thermo}, the role played by non-gravitational heating and cooling processes in altering the predictions of such a model is discussed; finally, we summarize the results presented in \S\ref{sec:conclusions}. Given the limited space available for this review, we refer the reader to recent reviews about early structure formation in the Universe \citep[e.g.][]{Bromm11} and cosmology with clusters of galaxies \citep[e.g.][]{Allen2011, kravtsov_borgani12} for a more extensive discussion of these topics. %---------------------------------------------------------% | \label{sec:conclusions} %---------------------------------------------------------% In this review, we have discussed recent results on structure formation focusing our attention on the first objects in the Universe and the most massive clusters of galaxies at the present day. These extreme scenarios allow us to clearly illustrate the relevance of the physics of plasma on the formation of cosmic structures along a wide range of spatial and temporal scales. In the hierarchical paradigm of structure formation, the first objects are the building blocks of subsequent structure formation, leading to larger galaxies and galaxy clusters through accretion and merger events \citep[e.g.][]{Somerville12}. Despite the disparity of involved scales, a number of physical processes, such as radiative cooling, turbulence and feedback, appear to be common, suggesting them to be ubiquitous in the non-linear regime of cosmic structure formation. In the early Universe, the first objects are expected to form in halos with $10^5-10^8$~M$_\odot$ at redshift $10-30$ \citep[e.g.][]{Bromm09, Bromm11}. Here we distinguish the so-called minihalos with virial temperatures above $1000$~K from the atomic cooling halos with virial temperatures above $10^4$~K. Minihalos are the expected formation sites for the first primordial stars, with typical masses in the range of $10-100$~M$_\odot$ \citep{Abel02, Bromm02, Yoshida08, Clark11, Greif11, Hosokawa11, Turk12, Latif13a, Susa13}. Their formation is governed by the chemistry and cooling of molecular hydrogen, as well as additional processes such as turbulence \citep[e.g.][]{Turk12, Latif13a}, radiative feedback \citep[e.g.][]{Hosokawa11, Susa13} and magnetic fields \citep[e.g.][]{Tan04, Machida06, Sur10, Schober12, Sur12, Turk12}. The atomic cooling halos show a more complex evolution depending on their local conditions, in particular regarding their metallicity and dust content. In this review, we restricted ourselves to the formation of massive black holes in primordial halos (see Fig.~\ref{fig:SMBH_flowchart} for an illustrative summary), while a more general discussion is given by \citet{Bromm11}. In the presence of strong photodissociating backgrounds, H$_2$ formation is suppressed \citep{Omukai01, Machacek01, Johnson08, Schleicher10, Shang10, Latif11}, leading to a close-to-isothermal collapse regulated via atomic hydrogen lines. Recent numerical simulations confirm that massive central objects can indeed form, due to the high accretion rates of more than $1$~M$_\odot$~yr$^{-1}$ \citep{Latif13b}. In the presence of such accretion rates, feedback can be expected to be weak \citep{Hosokawa12, Schleicher13} and does not impede the accretion. Indeed, even trace amounts of dust, corresponding to $10^{-5}-10^{-3}$ times the dust-to-gas ratio in the solar neighborhood, may already trigger strong cooling and fragmentation at high densities \citep{Schneider04, Omukai05}, but also stimulate the formation of molecular hydrogen at low to moderate densities \citep{Cazaux09, Latif12}. The extremely metal poor star SDSS J1029151+172927 \citep{Caffau11} shows chemical abundances at which metal line coolant is inefficient, and where only trace amounts of dust grains were able to trigger cooling and fragmentation \citep{Klessen12, Schneider12}\footnote{An alternative formation scenario for the star SDSS J1029151+172927 has been recently proposed by \citet{MacDonald_2013}, who suggest that it may have been a subgiant formed with significantly higher metallicity in the vicinity of a SN-Ia.}. For metallicities above $10^{-2}$ solar, on the other hand, metal line cooling can be expected to be significant \citep{Bromm03, Omukai05}. The fragmentation of such metal-enriched atomic cooling halos is in fact poorly understood \citep[see e.g.][for first modeling attempts]{Safranek10} and needs to be investigated in further detail. %-----------------------------------------------------------------------------------------------% \begin{figure} \centering\includegraphics[width=10cm, angle=0]{fig21_regan2009} \caption{Flowchart summarizing possible paths for the formation of the first SMBHs in high-redshift atomic cooling halos. Figure from \citet{Regan09a}.} \label{fig:SMBH_flowchart} \end{figure} %-----------------------------------------------------------------------------------------------% %-----------------------------------------------------------------------------------------------% \begin{figure} \centering %\psfig{file=fig20_pfrommer_2008_flowchart.eps,width=11.0truecm} \includegraphics[width=11.0truecm]{fig20_pfrommer_2008_flowchart} \caption{Flowchart summarizing the connections between the main physical processes taking place in galaxy clusters together with the different observational channels through which they can be detected. Figure from \citet{Pfrommer_2008}.} \label{fig:CR_flowchart} \end{figure} %-----------------------------------------------------------------------------------------------% In the hierarchical paradigm of structure formation, where the first objects are the building blocks of subsequent structure development, clusters of galaxies, with masses of up to $10^{15}M_\odot$ at $z=0$, occupy the most massive extreme of the cosmic hierarchy. The formation and evolution of galaxy clusters is a complex and non-linear event resulting from the intricate interaction of a number of physical processes acting on a wide range of scales \citep[see][for a recent review and references therein]{kravtsov_borgani12}. As an example, Fig.~\ref{fig:CR_flowchart} shows a simplified summary of some of the main processes operating in galaxy clusters. On large scales, the hierarchical process of structure formation induces the development of strong cosmological shocks, surrounding galaxy clusters and filaments, that contribute to heat and compress the hot intra--cluster plasma. Within galaxy clusters, weaker internal shocks, mainly originated by subhalo mergers or accretion phenomena, change the energetic balance of the gas and allow the halos to virialize. These shocks can also generate ICM turbulence and mixing, amplify magnetic fields, and accelerate thermal distributions of particles giving rise to a non-thermal population of CRs. In dense regions within galaxies and galaxy clusters, the intra--cluster gas can cool radiatively, leading to both star formation and gas accretion onto SMBHs residing at the center of massive cluster galaxies. These processes can then provide a significant energy contribution to the ICM in the form of SNe explosions or AGN feedback. As shown in Fig.~\ref{fig:CR_flowchart}, all these processes, which are highly interconnected between them, are manifested by means of different observational channels. In the last years, the new generations of supercomputing and programming facilities have been crucial to deepen in our understanding of the complicated physical processes taking place within the intra--cluster plasma and shaping the observational properties of galaxy clusters. In order to explain the observations, cosmological hydrodynamical simulations have tried to implement the most relevant physical processes self--consistently with the cosmic evolution. In particular, in addition to gravitationally--induced phenomena inherent to structure formation, the standard non--gravitational processes commonly included in these simulations are radiative cooling, star formation and SN feedback. In the last years, the inclusion of the effects of thermal and/or kinetic AGN feedback is also becoming a common practice, despite the fact that the particularities of the heating mechanism are still uncertain. In spite of the relatively simplicity employed in modeling these complex processes, simulations have been able to significantly reproduce most of the observational cluster properties, at least for massive systems at relatively outer cluster regions ($0.1R_{500} \mincir r \mincir R_{500}$), where clusters are assumed to be nearly self--similar. However, inner cluster regions and smaller systems show a number of significant issues that still need to be solved. In these inner regions, simulations still show an excess of gas cooling, which produces an excess in both the star formation and the metal production. In addition, simulations are still not able to solve the cooling flow problem or to reproduce the diversity of the observed temperature and entropy radial profiles of relaxed and unrelaxed systems. On the other hand, cluster outskirts ($r\magcir R_{500}$) are also affected by strong deviations from hydrostatic equilibrium caused, primarily, by sources of non-thermal pressure support such as CRs or magnetic fields, which generally are not modeled in simulations. % These results indicate that, in addition to the processes already included, a number of additional physical processes, mainly related with the complex physics of plasma, such as turbulence, viscosity or thermal conduction, must be also properly taken into account. Therefore, although AGN feedback seems to be the most favored energy source to regulate cooling in clusters, a subtle interplay with a number of supplementary physical phenomena may be needed to explain the observational properties of galaxy clusters and groups, from the core regions out to the outskirts. In the near future, a significant numerical effort will be aimed at performing larger and better--resolved cosmological simulations with a more accurate modeling of the physics of galaxy evolution. In addition to these technical improvements, forthcoming instruments, like the {\it JWST} \citep{Gardner_2006} and the new generation of large ground--based telescopes, are expected to detect light from the first galaxies, contributing to interpret early structure formation. Besides, a number of large observational surveys in different wavebands, such as {\it eROSITA} \citep{erosita_2012}, {\it Euclid} \citep{euclid_2012}, {\it WFXT} \citep{wfxt_2011} or the {\it LSST} \citep{lsst_2009}, will provide a significantly large number of clusters. These numerical and observational efforts, together with a more accurate treatment of the physics of plasma, will definitely shed some more light on the nature of the physical processes governing the formation of structures in the Universe, from the first non-linear objects to the present--day massive galaxy clusters. | 14 | 4 | 1404.3956 |
1404 | 1404.2932_arXiv.txt | The IceCube Neutrino Observatory has observed highly energetic neutrinos in excess of the expected atmospheric neutrino background. It is intriguing to consider the possibility that such events are probing fundamental physics beyond the standard model of particle physics. In this context, $\mathcal{O}$(PeV) dark matter particles decaying to neutrinos have been considered while dark matter annihilation has been dismissed invoking the unitarity bound as a limiting factor for the annihilation rate. However, the latter claim was done ignoring the contribution from dark matter substructure, which in a PeV Cold Dark Matter scenario, would extend down to a free streaming mass of $\mathcal{O}$($10^{-18}$M$_\odot$). Since the unitarity bound is less stringent at low velocities, ($\sigma_{\rm ann}$v)$\leq4\pi/m_\chi^2v$, then, it is possible that these cold and dense subhalos would contribute dominantly to a dark-matter-induced neutrino flux and easily account for the events observed by IceCube. A dark matter model where annihilations are enhanced by a Sommerfeld mechanism can naturally support such scenario. Interestingly, the spatial distribution of the events shows features that would be expected in a dark matter interpretation. Although not conclusive, 9 of the 37 events appear to be clustered around an extended region near the Galactic Center while 6 others spatially coincide, within the reported angular errors, with 5 of 26 Milky Way satellites. However, a simple estimate of the probability of the latter occurring by chance is $\sim35\%$. More events are needed to statistically test this hypothesis. PeV dark matter particles are massive enough that their abundance as standard thermal relics would overclose the Universe. This issue can be solved in alternative scenarios, for instance if the decay of new massive unstable particles generates significant entropy reheating the Universe to a slightly lower temperature than the freeze-out temperature, $T_{\rm RH} \lesssim T_{\rm f}\sim4\times10^4$~GeV. | The IceCube collaboration has recently announced the possible detection of the first high energy neutrinos with a cosmic origin \cite{IceCube_2013,IceCube_2013_2,IceCube_2014}. The all-sky search over a period of $\sim988$~days resulted in 37 events in the energy range between $\sim30$~TeV and $\sim2$~PeV. The possibility of these neutrinos having a purely atmospheric origin is currently ruled out at $\sim5.7\sigma$. Whether their origin is Galactic or extragalactic remains unknown with several {\it ordinary} astrophysical sources being considered so far (for an excellent review see \cite{Cosmic_PeV}). An intriguing possibility related to new physics is that of PeV dark matter decay or annihilation. A {\it smoking gun} dark-matter-induced monochromatic neutrino line might be consistent with both, an apparent drop-off feature above PeV energies in the neutrino spectrum, and the fact that the three highest neutrino events have similar energies ($1041^{+132}_{-144}$~TeV, $1141^{+143}_{-133}$~TeV and $2004^{+236}_{-262}$~TeV). The case of dark matter decay has been considered in detail elsewhere \cite{Feldstein_13,Esmaili_13,Bai_13,Bat_14} but dark matter annihilation has been discarded for the following reason: The rate of monochromatic neutrinos of energy $E_{\nu}$ produced by dark matter annihilation arriving at a detector on Earth of fiducial volume $V$ ($\sim1$~km$^3$ for IceCube) and nucleon number density $n_N$ ($\sim5\times10^{23}$cm$^{-3}$, the number density of ice) has been estimated as \cite{Feldstein_13}: \begin{equation}\label{base} \Gamma_{\rm Events}\sim V L_{\rm MW} n_N \sigma_N \left(\frac{\rho_\chi(R_\odot)}{m_\chi}\right)^2\left<\sigma_{\rm ann} v\right>\sim0.013~{\rm yr}^{-1}, \end{equation} using a neutrino-nucleon scattering cross section $\sigma_N\sim9\times10^{-34}$cm$^2$ at $E_\nu=m_\chi=1.2$~PeV \cite{Gandhi_98}. Dark matter is assumed to annihilate homogeneously across the characteristic length of the Milky Way (MW) galaxy $L_{\rm MW}\sim 10$~kpc, having a density equal to the estimated local value $\rho_\chi(R_\odot=8.5~{\rm kpc})=0.4$~GeVcm$^{-3}$ (consistent with current estimates, see e.g. \cite{Bovy_12}) and an annihilation cross section {\it exclusively into neutrinos} and saturated at the {\it local} unitarity limit: \begin{equation} \label{uni_local} \left<\sigma_{\rm ann} v\right>\equiv\frac{4\pi}{m_\chi^2\beta_{\rm loc}}, \end{equation} where $\beta_{\rm loc}\equiv v_{\rm loc}/c\sim10^{-3}$ is the typical {\it local} relative velocity of dark matter particles. With the estimate in Eq.~(\ref{base}), it seems that annihilation cannot account for the observed number of events. However, a more detailed calculation of the neutrino flux coming from {\it all} our MW halo should consider the following: (i) the change in dark matter density along the line of sight due to the radial dependence of the smooth dark matter distribution, which is enhanced towards the Galactic Centre; (ii) the contribution from dark matter substructure, and, more importantly, (iii) the unitarity limit depends on the relative velocity between dark matter particles at a given position along the line of sight. Thus, in principle, without violating the unitarity bound, the annihilation cross section could be much larger in the cold substructures present in our halo than at the solar circle as assumed in Eq.~\ref{base}. This type of behavior is natural in Sommerfeld-enhanced models, where $(\sigma_{\rm ann} v)\propto1/\beta$ is a common feature due the presence of a new mediator acting between the annihilating particles (e.g. \cite{Hisano_04,Arkani_09,Lattanzi_09}). In this paper we consider in detail (i)-(iii) to compute the rate of neutrino events potentially observable by IceCube and produced by PeV dark matter annihilation in our Galactic halo. The paper is organized as follows: In Section \ref{sec_one}, we describe how we estimate the contributions from the smooth dark matter distribution and from substructure (see also Appendix). We study the cases of a constant $(\sigma_{\rm ann} v)$ and one where $(\sigma_{\rm ann} v)\propto1/\beta$. The expected neutrino rate is presented in Section \ref{sec_results} as well as some indications of the compatibility of the spatial event distribution in the sky with that expected in a dark matter annihilation scenario. In Section \ref{sec_four}, the possible origin of PeV dark matter particles and their associated minimum self-bound halo mass are discussed. Finally we present a summary and our conclusions in Section \ref{conclusions}. | The announcement by the IceCube collaboration of the detection of over thirty neutrino events with a likely cosmic origin has been received with excitement raising a significant interest in discovering the responsible sources. At this moment, {\it ordinary} astrophysical sources (Galactic and/or extragalactic) could be responsible for the signal (for a review see \cite{Cosmic_PeV}), but the possibility of a dark matter origin is intriguing due to the connection with new physics beyond the Standard Model. The decay of PeV dark matter particles into neutrinos has been proposed recently \cite{Feldstein_13,Esmaili_13,Bai_13,Bat_14} while the case of PeV annihilating particles was dismissed invoking the unitarity bound to the annihilation cross section. In this work, we have revised the latter claim and compute in greater detail the expected rate of monochromatic neutrinos from PeV dark matter annihilation. We find that the unitarity limit can be satisfied and still produce sufficient PeV neutrinos if the cross section is enhanced by a Sommerfeld mechanism. In the simple case of $(\sigma_{\rm ann} v)\propto1/v$, the unitarity bound allows for a larger annihilation cross section in the cold subhalos, present in our Galactic halo, with a mass hierarchy going all the way down to the damping mass limit of PeV dark matter, $m_{\rm min}$ of $\mathcal{O}$($10^{-18}$M$_\odot$). In this scenario, to obtain the observed PeV neutrino rate, it is sufficient to saturate the cross section at a value of $\mathcal{O}$(100) times the {\it local} unitarity limit: $\left<\sigma_{\rm ann} v\right>_{\rm sat}\sim2.7\times10^{-23}$cm$^3$/s for $m_\chi=1$~PeV, at typical particle velocities of $\mathcal{O}$(1 km/s). A lower value of the {\it local} cross section would of course require a proportionally lower saturation velocity. The prediction in this model would be a signal with two main components: (i) a smooth dark matter contribution strongly peaked towards the Galactic Centre and (ii) and almost angle-independent contribution from dark matter subhalos where nearby large subhalos could appear as point sources. The relative contribution of these components across all angles would depend on the precise value of the velocity where the enhancement saturates. A lower saturation velocity would result in a stronger dominion of the subhalos. By looking at the all-sky distribution of events, we can see that a fraction of them ($\sim24\%$) are clustered around the Galactic Centre, although not at a strongly statistically significant level as pointed out before \cite{IceCube_2013}. Interestingly, 6 of the remaining 25 events coincide, within the angular errors, with the locations of five of the 26 MW satellites: Hercules, Sculptor, Sextans, Segue 1 and Ursa Major II. Although we have estimated that the probability of this (or more associated events) occurring randomly is $\sim35\%$, it would be worthy to test this possibility further once more events are collected. Regarding the origin of PeV annihilating dark matter particles, we have also revised the possibility of being produced as thermal relics of the Big Bang. Although strong constraints on very heavy dark matter relics have been derived in the past, they have so far ignored the substantial reduction of the relic abundance due to the Sommerfeld mechanism after kinetic decoupling. In the extreme case where $T_{\rm kd}\sim T_f$, the standard relic abundance for a constant s-wave annihilation gets suppressed by a factor of $0.2~{\rm ln}^{-1}(1/\sigma_{\rm vel}({\rm sat}))$, where $\sigma_{\rm vel}({\rm sat})$ is the 1D velocity dispersion of the dark matter particles when the Sommerfeld enhancement saturates. Even in this case however, the thermal relic abundance of PeV dark matter particles would overclose the Universe: $\Omega_\chi h^2\sim0.47$. A non-standard mechanism for dark matter production is therefore needed. For instance, if the Universe is reheated to a temperature $T_{\rm RH}\lesssim T_f$ by the decays of other unstable massive particles (e.g. moduli), then the relic abundance would be diluted by a factor of $(T_{\rm RH}/T_f)^3$. In this paper we have considered only the case of annihilation {\it exclusively} to neutrinos $\chi\chi\rightarrow\nu{\bar \nu}$, i.e., at tree-level, there is only a coupling to neutrinos. This produces a monochromatic neutrino signal. In a broader scenario, annihilation into other channels would lead also to a continuum of lower energy cosmic ray neutrinos (from the decay of the primary annihilation byproducts). This could in principle explain the more numerous sub-PeV events reported by IceCube. The combination of a monochromatic line with a continuum might even explain the gap feature that exists in the observed spectra between $\sim0.4$~PeV and $\sim1$~PeV. For the case of DM decay, this has been shown explicitly (see e.g. Fig. 6 of \cite{Bai_13}). It would be interesting to study particle physics models with the required spectra and yield in the case of dark matter annihilation. We note that a possible neutrino continuum could be obtained by the electroweak radiative corrections to the $\chi\chi\rightarrow\nu{\bar \nu}$ process. This possibility was studied in \cite{Serpico_07} where it was noted that the dominant $2\rightarrow3$ process is $\chi\chi\rightarrow\nu{\bar \nu}Z$. The authors estimated a branching ratio for this channel, $R=\sigma(\chi\chi\rightarrow\nu{\bar \nu}Z)/\sigma(\chi\chi\rightarrow\nu{\bar \nu})$, of $\mathcal{O}$(0.1) for $m_\chi\sim1$~PeV. Thus, a detailed analysis of this case might result in a non-negligible neutrino continuum. However, once other byproducts of the annihilation are considered, it is important to keep in mind current astrophysical constraints. For instance, in the case considered above, there should be an associated diffuse gamma-ray signal from neutral pion decay produced by quark jets from $Z$ decays \cite{Serpico_07}. The spectra $E^2dN/dE$ of these gamma-rays however, would peak at energies probably too high to put any significant constraint with current experiments, $E_{\rm peak}\sim m_\chi/30\sim33$~TeV\footnote{We obtain this approximate value by noting that for annihilation into quark-antiquark pairs, or $W$ and $Z$ bosons, the continuous gamma-ray yield is approximated by the following formula: $dN/dE\sim(0.42/m_\chi){\rm exp}[-8x]/(x^{3/2}+1.4\times10^{-4})$, where $x=m_\chi/E$ (e.g. \cite{Bergstrom_01}).}. For instance, the stringent current gamma-ray constraints for dark matter annihilation from observations of the MW dwarf spheroidals (dSphs) stand at \cite{Fermi_13}: \begin{equation}\label{const_fermi} \left<\sigma_{ \rm ann} v\right>(\chi\chi\rightarrow b{\bar b}, m_\chi=10~{\rm TeV})<\mathcal{O}(10^{-23}{\rm cm}^3{\rm s}^{-1}). \end{equation} At higher dark matter masses, there are no constraints, but we note that the model we have considered here would even be consistent at the level of Eq.~(\ref{const_fermi}). This is because dSphs have typical velocities of $\mathcal{O}$(10 km/s), and thus, $\left<\sigma_{\rm ann} v\right>_{\rm dSphs}\sim2.7\times10^{-24}$cm$^3$/s in the example we explored in this paper. The associated gamma-rays from dark matter annihilation in extragalactic halos are attenuated by the opacity of the Universe caused by pair production with the Extragalactic Background Light and the Cosmic Microwave Background radiation. The resulting electron-positron pairs loose energy via Inverse Compton scattering with the photon backgrounds. The final result is a cascade of the original high energy photons to lower GeV-TeV energies. This cascade is constrained by the extragalactic gamma-ray background observed by the {\it Fermi-LAT} instrument \cite{Abdo_2010}. In this way, it is possible to set an upper limit to the annihilation cross section in the channels that give rise to the original gamma-ray emission (e.g. \cite{Murase_2012}). Assuming a constant $\left<\sigma_{\rm ann} v\right>$ and a substructure boost more generous than the one we assumed here, this constraint stands at (see Fig. 15 of \cite{Murase_2012}): \begin{equation}\label{const_cascade} \left<\sigma_{ \rm ann} v\right>(\chi\chi\rightarrow [b{\bar b}~{\rm or}~W^{+}W^{-}~{\rm or}~\mu^+\mu^-], m_\chi=1~{\rm PeV})\lesssim3\times10^{-21}{\rm cm}^3{\rm s}^{-1}. \end{equation} This limit is satisfied by the maximum saturated cross section of the Sommerfeld-enhanced case studied here: $\left<\sigma_{\rm ann} v\right>_{\rm sat}\sim2.7\times10^{-23}$cm$^3$s$^{-1}$, which is two orders of magnitude lower. Another potential worry would be the energy injection in the early Universe due to dark matter annihilation. This could create distortions in the energy and power spectra of the CMB (e.g. \cite{Zavala_10,Slatyer_09}). The latter is the most constraining but still too weak at PeV masses to be of concern. The most recent analysis puts the following constraint \cite{Slatyer_13}: \begin{equation}\label{constraint} p_{\rm ann}=\frac{f_{\rm eff}\left<\sigma_{\rm ann} v\right>}{m_\chi}<1.18\times10^{-27}{\rm cm}^3{\rm s}^{-1}{\rm GeV}^{-1}, \end{equation} where $f_{\rm eff}$ is the efficiency factor to which the annihilation products get absorbed by the CMB plasma. For the case of a dominant channel of annihilation into neutrinos, most of the energy is lost and $f_{\rm eff}\ll1$. But even if one were to consider other annihilation channels and $f_{\rm eff}\sim1$, the constraint in Eq.~(\ref{constraint}) would be too weak for $m_\chi\sim1$~PeV. | 14 | 4 | 1404.2932 |
1404 | 1404.3942.txt | In this paper we present the most up-to-date list of nearby galaxies with optically detected supernova remnants (SNRs). We discuss the contribution of the H$\alpha$ flux from the SNRs to the total H$\alpha$ flux and its influence on derived star formation rate (SFR) {for 18 galaxies in our sample}. We found that the contribution of SNRs' flux to the total H$\alpha$ flux is $5\pm 5$\%. Due to the observational selection effects, the SNRs contamination of SFRs derived herein represents only a lower limit. | Understanding and modeling the star formation rates (SFRs) are the central goals of the theory of star formation. SFR is one of the crucial ingredients of cosmological simulations of galaxy formation, and these simulations demonstrate the impact of SFR models on galaxy evolution. The SFR has been a major issue in astrophysics since the 1970s, and during the last two decades, the cosmic star formation history of the universe has been widely studied in order to better constrain galaxy formation and evolution models.\footnote{Some recent models, however, show that galaxies self-regulate -- star formation is regulated by stellar feedback (radiation, stellar winds and supernovae) limiting the amount of very dense gas available for forming stars (Hopkins, Quataert \& Murray 2011).} Large-area surveys and use of larger telescopes have improved our knowledge of SFRs at low to intermediate redshifts ($z<1.0$; see Hopkins \& Beacom 2006) and beyond that (e.g. Karim et al. 2011, Sobral et al. 2013). These studies show strong decrease of SFR toward present epoch, and they suggest that the star formation activity over the last $\sim$11 Gyrs is responsible for producing $\sim$ 95 per cent of the total stellar mass density observed locally. Determination of SFRs in galaxies through the Hubble sequence provides vital clues to the evolutionary histories of galaxies. Measured SFRs are spread along six orders of magnitude (when normalized by galaxy mass), from almost zero in gas-poor elliptical, disk and dwarf galaxies, up to $\sim$100 M$_{\odot} \textrm{yr}^{-1}$, in optically-selected starburst galaxies, or even more in the most luminous infrared starburst galaxies (Kennicutt 1998). The first quantitative SFRs were derived from evolutionary synthesis models of galaxy colours (Searle, Sargent \& Bagnuolo 1973). Further on, more precise diagnostic tools were made using integrated emission-line fluxes (Kennicutt 1983), near-ultraviolet continuum fluxes (Donas \& Deharveng 1984), and infrared continuum fluxes (Rieke \& Lebofsky 1978). Nowadays, many different properties are used as star formation tracers, with the goal of directly or indirectly targeting continuum or line emission that is sensitive to the short-lived massive stars. Also, different techniques are used depending on whether we measure SFRs for the whole galaxy, or for regions within a galaxy (e.g. molecular clouds). The most common approach for measuring SFRs in resolved regions is to count individual objects (e.g. young stellar objects) or events (e.g. supernovae) that trace the recent star formation. Calibrations of SFR indicators have been made from the X-rays, ultraviolet (UV), via the optical and infrared (IR), all the way to the radio waves, using both continuum and line emission. For the most recent review on the latest achievements in the field see Kennicutt \& Evans (2012), as well as the previous review by Kennicutt (1998). In this paper, we focus on better constraining SFRs from the H$\alpha$ emission line. The next section will introduce star formation measurements from this emission line. \subsection{Star Formation Rates from H$\alpha$ flux} The nebular emission lines are very effective in re-emitting stellar luminosity, and thus provide direct measurement of young stellar content in the galaxy. This is especially the case for H$\alpha$ line, observationally the strongest among the recombination lines from regions of ionized gas surrounding young hot stars ({\hbox{H\,{\sc ii}}} regions). Only stars with masses exceeding 10 Solar masses contribute significantly to the stellar flux which can ionize interstellar medium (ISM). Also, these stars have a lifetime shorter then 20 Myr, so the emission lines give us almost instantaneous SFRs, independent of previous star formation histories. The conversion factor between integrated emission-line luminosity and SFR is computed using stellar evolutionary synthesis models. While calibrations have been published by numerous authors, here we use the calibration from Kennicutt, Tamblyn \& Congdon (1994), which assumes solar abundances and the Salpeter initial mass function (IMF) (Salpeter 1955) for stellar masses in range 0.1-100M$_{\sun}$: \begin{equation} \\\mathrm{SFR}\ (M_{\sun}\mathrm{yr}^{-1})=7.94\times10^{-42}L_{\mathrm{H}\alpha}\ (\mathrm{erg\ s^{-1}}), \end{equation} \noindent where calibration coefficient yields for the Case B recombination at $T_{\rm{e}}=10000K$. There is a significant variation among published values for this calibration coefficient ($\sim30\%$). The differences in coefficient values originate in the usage of different stellar evolution, atmosphere models and IMFs. The method for deriving SFRs from the H$\alpha$ flux is one of the most used methods due to the advantages of optical observations and strength of the H$\alpha$ line. The star formation in nearby galaxies can be mapped at high resolution even with small telescopes, and the H$\alpha$ line can be detected in the redshifted spectra of starburst galaxies up to $z<3$ (Bechtold et al. 1997, Geach et al. 2008, Sobral et al. 2013). The main limitations of the method are its sensitivity to uncertainties in extinction and the IMF, and the assumption that all of the massive star formation is traced by the ionized gas. There is a fraction between 15\% - 50\%, of ionizing radiation which escapes from {\hbox{H\,{\sc ii}}} regions. This can be measured either directly (Oey \& Kennicutt 1997), or from the observations of the diffuse H$\alpha$ emission in nearby galaxies. In their analysis of diffuse H$\alpha$ emission Ferguson et al. (1996) found that diffuse ionized gas is photoionized by Lyman continuum photons which have escaped from {\hbox{H\,{\sc ii}}} regions. Therefore, diffuse H$\alpha$ emission should be included when this SFR measurement method is used. The dominant source of systematic error in SFR measurements from H$\alpha$ fluxes is extinction within the galaxy observed. Internal extinction can be corrected by combining H$\alpha$ line and IR continuum emission or radio data {(Kennicutt et al. 2009)}, which are not affected by the extinction. Mean extinction values range from A(H$\alpha$)=0.5 mag to A(H$\alpha$)=1.8 mag (see Kennicutt (1998) and James et al. (2005) and references therein), depending on the galaxy type and luminosity. As this extinction correction comes with large uncertainty, and therefore represents a large change (2 - 3 times higher) in derived fluxes, this correction {was usually} not applied when calculating SFRs. {With recent IR surveys (\textit{Spitzer}, \textit{Herschel}, \textit{WISE}), which give estimates on dust amount in nearby galaxies, extinction correction is more commonly considered.} Also, we emphasize the importance of eliminating H$\alpha$ flux contaminants when calculating SFRs from H$\alpha$ emission. Emission spectra from {\hbox{H\,{\sc ii}}} regions show that very close to the H$\alpha$ line, on both sides, are \hbox{[N\,{\sc ii}]} lines at $\lambda$654.8 nm and $\lambda$658.3 nm. Most of the filters used to extract H$\alpha$ line also let some of the \hbox{[N\,{\sc ii}]} emission pass through. Using different methods, this problem can be minimized. However, it leaves some uncertainty in derived H$\alpha$ fluxes. Using spectroscopic observations, we can calculate the ratio between H$\alpha$ and \hbox{[N\,{\sc ii}]} lines, and using filter profiles, we can obtain corrections for \hbox{[N\,{\sc ii}]} contamination. Another possibility is to use very narrow \hbox{[N\,{\sc ii}]} filters for deriving this correction (James et al. 2005). However, because of gradient in \hbox{[N\,{\sc ii}]} abundances with change of galactocentric distance, one should be careful when applying any correction. The most commonly used corrections for \hbox{[N\,{\sc ii}]} emission for entire galaxies are those derived by Kennicutt (1983) and Kennicutt \& Kent (1983). Using Sloan Digital Sky Survey Data Release 4 galaxies (SDSS DR4, Adelman-McCarthy et al. 2006), a simple correction has also been derived by Villar et al. (2008) and presented by Sobral et al. (2012) and is now widely used. Another origin of H$\alpha$ flux contamination are the sources which emit H$\alpha$ radiation, but which are not {\hbox{H\,{\sc ii}}} regions surrounding young high-mass OB stars. There is a wide range of such objects which cause overestimate in SFRs: another emission nebulae - planetary nebulae (PNe) and SNRs; active galactic nuclei (AGNs); ultraluminous X-ray sources (ULXs) and their surrounding nebulae; {micro-quasars;} foreground stars with H$\alpha$ emission; superbubbles. None of these types of objects have so far been thoroughly discussed as possible sources of systematic error in H$\alpha$ flux-based determination of SFRs. In some papers, there are efforts to exclude emission from PNe when making catalogues of {\hbox{H\,{\sc ii}}} regions and estimating SFRs. Azimlu, Marciniak \& Barmby (2011) did this for M31 galaxy, and when they removed all PNe candidates, they found that PNe were responsible for 1\% of the total measured H$\alpha $ emission. The problem with PNe, as well as with SNRs, is that they are hard to differentiate from {compact} {\hbox{H\,{\sc ii}}} regions based on H$\alpha$ emission only. Additional observations in narrow band filters can certainly distinguish between different emission nebulae types. Also, PNe are much smaller in sizes, when compared to the majority of {\hbox{H\,{\sc ii}}} regions. HII regions can be compact, the same in size as PNe - smaller than a parsec, if they are excited only by a single star, but {\hbox{H\,{\sc ii}}} regions are generally larger, being excited by multiple young massive stars, or even clusters. On the other hand, SNRs can be also different in sizes, from being compact (few parsecs), to huge (up to the 100pc; {\v{C}ajko, Crawford \& Filipovi\'{c} (2009)}), depending on the phase of their evolution, as well as on input energy from SN explosion and surrounding ISM density. Some AGNs are prominent in H$\alpha$ line from their broad-line and narrow-line regions (it depends on the redshift and the inclination angle). But, considering their small size, and certainly a modest number of them which could be detected in the projection of a single galaxy, their contribution to the total H$\alpha$ flux of a galaxy would be negligible. On the other hand, there are efforts to exclude AGNs from a sample of star-forming galaxies at some higher redshift, which are studied for SFRs (Garn et al. 2010, Villar et al. 2011). ULX sources are compact X-ray sources that are located away from the nucleus of their host galaxy, emitting well above the Eddington limit of a 20M$_{\sun}$ black hole (L$_{X}$ $\sim3\times10^{39}$ erg s$^{-1}$). Recently, more and more ULX sources have been detected with strange nebular emission surrounding it. Some of them are IC342 X-1 source in IC342 galaxy ({Roberts et al. 2003, Abolmasov et al. 2007, Feng \& Kaaret 2008}), Holmberg IX X-1 in Holmberg IX galaxy ({Fabbiano 1988, Gladstone, Roberts, \& Done 2009, Moon et al. 2011, Gris\'{e} et al. 2011}) and MF 16 (nomenclature from Matonick \& Fesen 1997) in NGC6946 {(Roberts \& Colbert 2003)}. Most likely this nebular emission is coming from the accretion disks. Such nebulae are also prominent in H$\alpha$ line and, as such, are potential contaminants of H$\alpha$ flux and instantaneous SFRs. Of course, ULX sources can be regarded as tracers of some past SFRs. {In NGC6946 galaxy, flux of the MF16 is responsible for 0.1\% of the total H$\alpha$ flux, as is the case with IC342 galaxy and IC342 X-1 source. On the other hand, as Andjeli\'{c} (2011) has shown, H$\alpha$ derived SFRs for Holmberg IX galaxy can be significantly changed if nebular emission from the ULX is removed from the integrated H$\alpha$ flux of the galaxy. Emission from HoIX X-1 is resposible for even 75\% of the H$\alpha$ emission in this dwarf galaxy. We adopt H$\alpha$ flux of this object from Arbutina et al. (2009).} { Microquasars are X-ray binary systems, with accretion disk, relativistic jets, and usually with strong and variable radio emission. Also, they can have surrounding emission nebula. Large nebula S26 (from Blair \& Long 1997) in the nearby galaxy NGC7793 is a jet-inflated bubble around a powerful microquasar (Pakull, Soria \& Motch 2010). H$\alpha$ emission from this source is not tracer of a star formation, and it is responsible for 1.15\% of the galaxy's total H$\alpha$ emission.} Foreground emission-line stars, such as O and B supergiants and Wolf-Rayet stars exhibit strong optical emission lines, primarily hydrogen Balmer lines. In most of these systems, emission originates from strong stellar winds. While these stars have strong emission, they also have very bright continuum, and therefore they should be easily distinguished from {\hbox{H\,{\sc ii}}} regions. In that sense, foreground emission-line stars could be easily removed from the total H$\alpha$ flux and should not represent a source of error in H$\alpha$ derived SFRs. Superbubbles are regions of bright emission which frequently surround OB association. This kind of nebulae is powered by a combination of stellar winds, UV stellar radiation and occasional SN explosions. On the other hand, wind-blown nebulae are related to the {\hbox{H\,{\sc ii}}} regions, but they are ionized due to shock fronts caused by strong stellar winds. Superbubbles are characterized by low velocity shock fronts {($<$ 100 km s$^{-1}$)}, while classical SNRs have shock fronts expanding at 100 - 1000 km s$^{-1}$. With modern observational equipment, superbubbles can be distinguished from the SNRs, mostly upon detection of OB associations inside. {In M31 SNR survey, Lee \& Lee (2014) separated 44 superbubbles from previous SNR candidates, which represent 1.3\% of total H$\alpha$ emission of this galaxy, while in M101 galaxy 10 superbubbles represent 0.3\% (Franchetti et al. 2012).} %Superbubbles and diffuse ionized gas - Long at al. 2010 strana 20; Blair \& Long (1997) strana 16; Pannuti et al. 2002 strana 3. Franchetti et al. 2012. The main aim of this paper is to estimate the influence of SNRs' emission on the H$\alpha$ derived SFRs. In order to do so, we assemble current sample of optically detected SNRs in nearby galaxies. {In the next section we discuss optical detection of SNRs. Section 3 gives details on individual galaxies in our sample and SNR detections in those galaxies, while in section 4 we discuss influence of SNRs' emission on the H$\alpha$ derived SFRs. In section 5 we present our conclusions. } | {Out of 25 nearby galaxies with observed SNRs in optical range, we have presented details on SNRs detection for our sample of 18 galaxies.} We have discussed the contribution of the H$\alpha$ fluxes from the SNRs to the total H$\alpha$ flux and its influence to the derived SFR for each galaxy in the sample. We have found that the {average SNR contamination to the total H$\alpha$ flux and derived SFRs for analysed nearby galaxies is $5\pm 5$\%. The highest SNR contamination is about 13\%.} M83 galaxy is best sampled for optical SNRs and it has {9\%} of SNRs in total H$\alpha$ emission. We expect that percentages similar to this one should be close to the real contribution of the SNRs emission to the total H$\alpha$ emission in spiral galaxies. Due to the selection effects, the SNR H$\alpha$ contamination obtained in this paper represent only a lower limit. | 14 | 4 | 1404.3942 |
1404 | 1404.7504_arXiv.txt | We investigate pre-processing using the observed quenched fraction of group and cluster galaxies in the \citet{yang07} SDSS-DR7 group catalogue in the redshift range of $0.01 < z < 0.045$. We categorize group galaxies as virialized, infall or backsplash and we apply a combination of the Dressler-Shectman statistic and group member velocities to identify subhaloes. On average the fraction of galaxies that reside in subhaloes is a function of host halo mass, where more massive systems have a higher fraction of subhalo galaxies both in the overall galaxy and infall populations. Additionally, we find that between $2 \lesssim r_{\text{200}} < 3$ the quiescent fraction is higher in the subhalo population with respect to both the field and non-subhalo populations. At these large radii ($2 \lesssim r_{\text{200}} < 3$), the majority of galaxies ($\sim 80 \%$) belong to the infall population and therefore, we attribute the enhanced quenching to infalling subhalo galaxies, indicating that pre-processing has occurred in the subhalo population. We conclude that pre-processing plays a significant role in the observed quiescent fraction, but only for the most massive ($M_{\text{halo}} > 10^{14.5} M_{\odot}$) systems in our sample. | Observational studies of rich galaxy clusters have shown that most of the members are red early-type galaxies with little or no on-going star formation \citep{oemler74, dressler80, blanton03, balogh04, baldry06}. While a high fraction of quiescent (i.e.\ not actively star-forming) galaxies have been observed in rich groups and clusters \citep{kauffmann04, wilman05, peng10, mcgee11, muzzin12}, recent results from observations and simulations (both numerical and semi-analytic) indicate that star formation quenching actually begins in low mass haloes with $M_{\text{halo}} \sim 10^{13} M_{\odot}$ \citep{mcgee09,bm10,george11,delucia12,wetzel12}. Additionally, there is evidence that some cluster galaxies had their star formation quenched in groups with $M_{\text{halo}} \geq 10^{13} M_{\odot}$ prior to accretion onto the more massive cluster environment, a process often referred to as pre-processing \citep{zm98b,km08,berrier09,mcgee09,delucia12}. While quenching has been shown to occur in low mass haloes, the significance of pre-processing is still a subject of debate. Using $N$-body simulations \citet{berrier09} found that $70 \%$ of their cluster ($10^{14} < M_{\text{halo}} < 10^{14.6} M_{\odot}$) galaxies fell in directly from the field, while only $\sim10 \%$ fell in as members of group-sized haloes with $M_{\text{halo}} \geq 10^{13} M_{\odot}$. Based on these results, \citet{berrier09} concluded that pre-processing did not significantly contribute to the quenched fractions observed in present-day clusters. In contrast, both \citet{mcgee09} and \citet{delucia12} used semi-analytic models (SAMs) to show that $\sim25-45 \%$ of their simulated cluster galaxies fell in as members of systems with $M_{\text{halo}} \geq 10^{13} M_{\odot}$, where the range depends on the mass of the galaxy and the mass of the host cluster. It should be noted that according to \citet{delucia12}, part of the discrepancy between the results of \citet{berrier09} and \citet{mcgee09} arises from differing definitions of `satellite', with the former computing fractions based on the time when a galaxy \emph{first} becomes a satellite of any halo and the latter when a galaxy becomes a satellite of the \emph{final or present-day} group or cluster. With the former definition, \citet{delucia12} find that their results are not inconsistent with those of \citet{berrier09}. A similar analysis was carried out using $N$-body hydrodynamical simulations by \citet{bahe13}. These authors found that $\sim 15 - 60 \%$ of galaxies in host haloes in the mass range of $10^{13.5} < M_{\text{halo}} < 10^{15.2} M_{\odot}$ had been pre-processed where the amount of pre-processing scaled with halo mass; massive haloes had a higher fraction of pre-processed galaxies \citep{bahe13}. Thus, the results of some SAMs \citep[e.g.][]{mcgee09,delucia12} and numerical simulations \citep[e.g.][]{bahe13} predict that pre-processing can play an important role in quenching star formation, especially in massive clusters. If the simulation predictions
of significant pre-processing
in groups and
clusters are correct
then it should be possible to
observe pre-processing
by looking at the populations
of galaxies in different environments
. The aim of this paper is to investigate the significance of pre-processing in a statistical sample of observed groups and clusters. Pre-processing can be investigated by studying the properties of \emph{infalling} subhalo galaxies, where a subhalo is defined as a collection of galaxies that reside in a small halo embedded within a larger parent halo. Subhaloes can be identified by performing substructure analysis with the Dressler-Shectman (DS) Test \citep{ds88}, which can detect galaxies with kinematic properties that deviate from those of the host halo. It should be noted that this method of identifying subhaloes differs from those used in numerical simulations. In particular, our observational definition of subhaloes is based on identification of kinematically distinct galaxies and does not require the galaxies within the subhalo to be gravitationally bound to one another, which is usually the case for subhaloes identified in simulations. Subhaloes, detected via the DS Test, are preferentially found on the group or cluster outskirts \citep{wb90,zm98a,hou12,dressler13} and the usual assumption is that these systems are infalling. However, numerical simulations have shown that a large fraction of galaxies beyond the virial radius, and out to $\sim 2.5$ virial radii, have already passed through the group or cluster core \citep[i.e.\ backsplash galaxies:][]{balogh00, mamon04, gill05, mahajan11, pimbblet11,bahe13, oman13}. Backsplash galaxies may have experienced star formation quenching due to more massive group - or cluster-related processes, and it has been suggested that much of the environmental quenching beyond the virial radius (out to $\sim2.5$ virial radii) is most likely due to the presence of a backsplash population \citep{wetzel14}. In contrast, the infall population typically refers to galaxies that are infalling onto the host system for the first time. Thus, any observed enhanced quenching must be a result of a transformation that occurred prior to accretion onto the host halo. Currently, most methods of distinguishing between the virialized, infall and backsplash populations in observed groups and clusters are based on the results of simulated systems. These classification schemes typically involve examining $|\Delta cz|/\sigma$ distributions \citep{gill05, pimbblet11} or dividing the $\Delta cz/\sigma - r_{\text{200}}$ plane into regions occupied by virialized, infall and backsplash galaxies \citep{mahajan11, oman13}. Although each population resides in a distinct region in the full phase-space of simulated clusters, projection effects can distort these clear divisions and there is often contamination between the observed populations (to be discussed in more detail in Section \ref{popfracs}). In addition to differences in their phase-space locations, infall and backsplash galaxies should also have subtle differences in their stellar mass distributions. As a result of tidal disruption, backsplash galaxies should be on average less massive than infalling galaxies at the same radius \citep{gill05}, and galaxies infalling in subhaloes will typically be more massive than individual infalling galaxies \citep{mcgee09}. Thus, in order to better probe pre-processing and environmental effects on galaxy evolution, it is important to examine the properties of virialized, infall and backsplash galaxies as a function of stellar mass and over a wide range of masses. % Although it is well known that high-density environments, such as groups and clusters, show signs of enhanced star formation quenching with respect to the field \citep{kauffmann04, rines05, kimm09, wetzel12, woo13}, the processes that dominate this transformation are still debated. Comparing the properties of infalling and backsplash subhalo galaxies allows us to probe the relative importance of rich group- and cluster-related processes, which are observable in the backsplash population, to pre-processing in lower mass haloes, which can be observed in the infalling subhalo population. In this paper we use a well-studied SDSS group catalogue to probe the properties of subhalo galaxies in groups and clusters in order to investigate the amount of pre-processing that occurs and to study the relative importance of the lower mass group environment in the evolution of galaxies. The paper is structured as follows: in Section \ref{data_p3}, we present our group and galaxy sample and in Section \ref{identifying_sub}, we discuss how we identify subhaloes. We compare the properties of the non-subhalo and subhalo populations, as well as compare the virialized, infall and backsplash subpopulations in Section \ref{subgals}. Finally, in Section \ref{preprocess}, we discuss our results and present our conclusions in Section \ref{conclusions}. Throughout this paper we assume a $\Lambda$CDM cosmology with $\Omega_{m,0} = 0.31$, $\Omega_{\Lambda,0} = 0.69$ and $H_{0} = 70$ km s$^{-1}$ Mpc$^{-1}$. | \label{conclusions} We have looked at the infall and backsplash subhalo populations in SDSS-DR7 groups and clusters, using a sample of satellite galaxies, which is complete to $M_{\text{star}} = 3.16 \times 10^{9} M_{\odot}$. The aim of this work is to investigate the importance of pre-processing in group and cluster galaxies. The Dressler-Shectman Test was used to identify subhalo galaxies and we followed the methodology of \citet{mahajan11} to classify virialized, infall and backsplash galaxies. The main results of this analysis are: \begin{enumerate} \item Subhaloes preferentially reside in massive systems and at large group-centric radii; \item The stellar mass distributions of non-subhalo and subhalo galaxies are marginally distinct, where subhaloes have, on average, slightly more massive galaxies; \item Low and intermediate mass group galaxies out to $3 r_{\text{200}}$ and high mass satellites close to the group core show enhanced SF quenching with respect to the field; \item On the group and cluster outskirts, between $2 \lesssim r_{\text{200}} < 3.0$, \fq is higher in galaxies that reside in subhaloes than for the overall satellite galaxy population at all stellar masses; \item As a function of radius, the percentages of infall and backsplash galaxies do not differ between non-subhalo and subhalo galaxies; \item Below halo masses of $\sim 10^{13.2} M_{\odot}$, all groups do not contain any detected subhaloes, while more massive haloes show a scatter in the values in both the fraction of subhalo galaxies ($f_{\text{sub}}$) and the fraction of infalling galaxies in subhaloes ($f_{\text{sub, infall}}$). Additionally, there appears to be a trend between $f_{\text{sub}}$ and $f_{\text{sub, infall}}$ with halo mass, where more massive haloes have both more subhaloes and more infalling subhaloes. \end{enumerate} The observed enhanced quenching in infalling subhalo galaxies, defined based as having kinematic properties distinct from the host group, suggests that pre-processing does play a role in galaxy evolution; however, the significance of pre-processing depends on halo mass. Pre-processing does not appear to be the dominant mechanism in groups and low mass clusters ($M_{\text{halo}} \lesssim 10 ^{14.5} M_{\odot}$), but it does play a significant role in producing the observed quenched fraction in \emph{massive clusters} with $M_{\text{halo}} > 10^{14.5} M_{\odot}$. | 14 | 4 | 1404.7504 |
1404 | 1404.7218_arXiv.txt | { LAMOST has released more than two million spectra, which provide the opportunity to search for double-peaked narrow emission line (NEL) galaxies and AGNs. The double-peaked narrow-line profiles can be well modeled by two velocity components, respectively blueshifted and redshifted with respect to the systemic recession velocity. This paper presents 20 double-peaked NEL galaxies and AGNs found from LAMOST DR1 using a search method based on multi-gaussian fit of the narrow emission lines. Among them, 10 have already been published by other authors, either listed as genuine double-peaked NEL objects or as asymmetric NEL objects, the remaining 10 being first discoveries. We discuss some possible origins for double-peaked narrow-line features, as interaction between jet and narrow line regions, interaction with companion galaxies and black hole binaries. Spatially resolved optical imaging and/or follow-up observations in other spectral bands are needed to further discuss the physical mechanisms at work. | % \label{sect:intro} The search for double-peaked narrow-line structure in galaxy spectra is an effective way of finding binary AGN candidates (Zhou et al. \cite{ref57}; Blecha et al. \cite{ref1}), which are expected to take place in the final phases of the merging of two interacting active galaxies. Since the suggestion that double-peaked profiles could be produced by binary AGN (Zhou et al. \cite{ref57}), much attention has been paid to that hypothesis, yet few definite cases are known. The most convincing examples can be found in CXO J1426+35 (Barrows et al. \cite{ref25}), EGSD2 J1420+4259 (Gerke et al. \cite{ref2}), NGC6240 (Komossa et al. \cite{ref59}), COSMOS J100043+020637 (Comerford et al. \cite{ref21}), SDSS J0952+2552 (McGurk et al. \cite{ref60}; Fu et al. \cite{ref26}) and so on (see the review in Wang et al. \cite{ref58}, and Table 1 in Ge et al. \cite{ref9} ). Several systematic searches for AGNs with double-peaked \oiii\ emission lines have been performed in the DEEP2 survey sample (Gerke et al. \cite{ref2}; Comerford et al. \cite{ref3}) and in the Sloan Digital Sky Survey (hereafter SDSS) data releases (Xu \& Komossa \cite{ref4}; Wang et al. \cite{ref5}; Liu et al. \cite{ref6, ref7}; Smith et al. \cite{ref8}; Ge et al. \cite{ref9}). The last authors have retrieved a large sample of 3030 dual-peak NEL objects from the SDSS DR7, and a much larger sample of asymmetric profiles NEL objects, with in total 54 dual-cores candidates. Among the candidates found at low or moderate redshift, when double-peaked \oiii\ emission line has been suspected or confirmed as indicator of a binary AGN, the components most likely trace objects with spatial separations in the range from 100 pc to 10 kpc (Wang et al. \cite{ref5}). However, the double-peaked NEL may also be produced by other mechanisms, such as chance superposition, peculiar gas kinematics in the narrow-line regions (NLRs; e.g., Gelderman \& Whittle \cite{ref32}; Fu \& Stockton \cite{ref33}; Fischer et al. \cite{ref34}; Shen et al. \cite{ref24}; Fu et al. \cite{ref26}) and jet-cloud interactions (e.g., Stockton et al. \cite{ref35}; Rosario et al. \cite{ref36}). Because of the growing interest in double-peaked NEL objects, it is also important to check that the candidates are fully trustworthy, thus independant surveys should confirm this character. In order to further analyse these double-peaked samples, follow-up observations are needed. Previous works have used various methods: high-resolution optical imaging (Comerford et al. \cite{ref21}), near-infrared imaging (Liu et al. \cite{ref6}; Fu et al. \cite{ref22}; Rosario et al. \cite{ref23}; Shen et al. \cite{ref24}; Barrows et al. \cite{ref25}), integral-field spectroscopy (Fu et al. \cite{ref26}), hard X-ray observations (Comerford et al. \cite{ref27}; Civano et al. \cite{ref28}; Liu et al. \cite{ref29}), radio observations (Zhou et al. \cite{ref57}; Fu et al. \cite{ref30}), and long slit spectroscopy (Shen et al. \cite{ref24}; Comerford et al. \cite{ref31}). In this paper, we focus on a systematic search for galaxies and AGNs with double-peaked NEL in the First Data Release of the Sky Survey conducted with the Large Sky Area Multi-Object Fiber Spectroscopic Telescope (LAMOST, also called Guo Shou Jing Telescope, GSJT), hereafter LAMOST DR1. Like the SDSS, LAMOST has produced a large number of quasar and galaxy spectra from October 2011 till June 2013 (its first two observing seasons). We have developed a method for searching double-peaked NEL galaxies and AGNs in LAMOST DR1, and visually inspected the candidates spectra for confirmation. Our sample selection and method are described in Section 2. We present and discuss the results in Section 3, and compare them to SDSS spectra derived results when available. We summarize this work in Section 4. | \subsection{Optical data} We find a total of 20 double-peaked NEL objects, 15 galaxies and 5 QSO, listed in Table \ref{tab1-list}. Among them, 10 have already been published by other authors, the remaining 10 being first discoveries. Three new spectra were first observed by LAMOST, see the Fig. \ref{fig3-spectra}. All our candidates have a high S/N with an average r-band magnitude (in SDSS system) of $\overline{r}$ = 17.17. \begin{table}[htb!] \caption[]{List of double-peaked NEL sample in LAMOST DR1} \label{tab1-list} \centering \setlength{\tabcolsep}{2pt} \begin{tabular}{cccccccccc} \hline \hline\noalign{\smallskip} Designation & Obs.Date & MJD & Plate & SPID & FIB & Type & z & r & SDSS\\ \scriptsize(1) & \scriptsize(2) & \scriptsize(3) & \scriptsize(4) & \scriptsize(5) & \scriptsize(6) & \scriptsize(7) & \scriptsize(8) & \scriptsize(9)\\ \hline\noalign{\smallskip} J005407.03+073611.9 & 20121108 & 56240 & EG010249N073002F & 14 & 047 & G & 0.10719 & 17.05 & 0\\ J023658.06+024217.9 & 20121114 (*)& 56246 & EG023131N032619F & 06 & 232 & G & 0.08744 & 16.17 & 1\\ J023832.67+023349.1 & 20121114 & 56246 & EG023131N032619F & 06 & 206 & G & 0.20780 & 17.13 & 0\\ J083425.28+283451.3 & 20130113 & 56306 & HD083217N291909F01 & 08 & 157 & G & 0.10253 & 17.04 & 1\\ J091646.03+283526.7 & 20120201 & 55959 & B5595902 & 15 & 006 & G & 0.14233 & 16.63 & 1\\ J094430.79+435421.4 & 20130215 & 56339 & HD094642N450651F01 & 05 & 088 & Q & 0.58779 & 17.77 & 1\\ J100708.01+242039.0 & 20130401 & 56384 & HD100153N235852F01 & 09 & 134 & Q & 0.54355 & 18.81 & 1\\ J104718.31+254348.3 & 20130210 & 56334 & HD104049N254200F01 & 09 & 237 & Q & 0.62712 & 18.72 & 1\\ J113630.61+135848.8 & 20130214 & 56338 & HD112941N152447F01 & 07 & 088 & G & 0.08169 & 16.91 & 1\\ J121342.90+422202.8 & 20130307 (**) & 56359 & HD121906N401846F01 & 16 & 245 & G & 0.07525 & 15.46 & 1\\ J123314.49+262624.9 & 20130111 & 56304 & HD122624N271605F01 & 06 & 006 & Q & 0.57025 & 18.58 & 1\\ J131434.73+563419.3 & 20130430 & 56413 & HD132545N565813F01 & 10 & 239 & G & 0.14407 & 17.39 & 1\\ J133730.29-002525.4 & 20130406 & 56389 & HD134427N004207F01 & 10 & 050 & G & 0.17212 & 17.11 & 1\\ J133737.82+555816.7 & 20130430 & 56413 & HD132545N565813F01 & 06 & 023 & Q & 0.47461 & 19.26 & 1\\ J135207.73+052555.8 & 20130409 & 56392 & HD135024N052949M01 & 04 & 239 & G & 0.07892 & 15.01 & 1\\ J135646.10+102609.0 & 20120514 & 56062 & VB3\_210N09\_V2 & 14 & 091 & G & 0.12313 & 15.81 & 1\\ J140225.68+465817.4 & 20130310 & 56362 & HD141351N461930F01 & 14 & 218 & G & 0.12759 & 16.79 & 1\\ J140231.28+184807.8 & 20130210 & 56334 & HD140137N164527F01 & 11 & 118 & G & 0.13970 & 18.95 & 0\\ J150501.56+371311.7 & 20130402 & 56385 & HD145553N362548F01 & 13 & 109 & G & 0.06517 & 16.25 & 1\\ J232703.17+004256.7 & 20121024 & 56225 & EG232111N021150M01 & 07 & 073 & G & 0.06601 & 16.47 & 1\\ \noalign{\smallskip} \hline \end{tabular} \tablecomments{0.86\textwidth}{ Column 1: LAMOST DR1 designation hhmmss.ss+ddmmss.s (J2000.0); Column 2: Date of observation; Columns 3-6: the MJD, Plate, Spectrograph id and Fiber id in LAMOST Sky Survey; Column 7: Type: G: galaxy nucleus, Q: QSO; Column 8: redshift; Column 9: apparent magnitude in r-band (in SDSS system); Column 10: 1 or 0 means that the object has a spectrum in SDSS DR9 or not.\\ (*) observed also on 20130501, (**) observed also on 20121113 } \end{table} Fig. \ref{fig4-dis}a shows their redshift distribution, which is strongly bimodal because of our selection criteria. Objects with $0.5 \leq z \leq 0.7$ are all optically unresolved QSO, for which the emission complex \ha\ - \nii\ is shifted beyond the available spectral range. The double-peaked NEL character for these objects is only determined from the profiles of \oiii $\lambda\lambda\ 4959, 5007$ and the narrow component of \hb. An additional constraint is added to the initial trimming criteria (section 2.2.1 above) because of poor quality of LAMOST spectra beyond 8500 $\AA$, as a redshift cut at $z = 0.7$ so that \oiii $\lambda 5007$ is measurable in good conditions. Objects with $z \leq 0.2$ are resolved galaxies on the SDSS images. To ensure a reliable identification of double NEL, we require that not only \hb\ and \oiii\ $\lambda\lambda\ 4959, 5007$, but also \nii $\lambda\lambda\ 6548, 6584$, \ha\ and \sii $\lambda\lambda\ 6717, 6731$ all present double-peaked narrow-line profiles, and for these we require as additional constraint that \sii $\lambda\ 6731$ $<$ 8500$\AA$. Note that this difference in selection criteria between galaxies and QSO is motivated by the data quality: in LAMOST DR1, the bulk of the redshift distribution of QSOs is clearly higher than that of the galaxies (at a given average magnitude) and the signal-to-noise of QSO spectra is generally better than that of galaxy nuclei, probably because of illumination conditions of the entrance face of the fiber. Fig. \ref{fig4-dis}b shows the distribution of the velocity differences between the two components of the NEL. The distribution is roughly symmetric around a mean value of $ <\Delta V>$ = 280 \kms . There is no significant difference in the observed $ <\Delta V>$ between QSO and galaxies. The spectral resolution of LAMOST corresponds to an average instrumental FWHM of 3.5 $A$, i.e. 200 \kms on the \oiii $\lambda 5007$ at the average redshift of the galaxy subsample (for higher z or lines in the red range the instrumental constraints are less stringent). The accuracy of centering of a good s/n gaussian component having such an FWHM, with usual computer algorithms, is roughly 1/10 of the FWHM, degrading when the s/n weakens. Hence we are confident that all our detected double peaked NEL listed in Table \ref{tab2-emission} are reliable. Further, the same procedures applied to the 17 objects that have a spectrum in the SDSS DR9 yielded consistent results with those listed in Table \ref{tab2-emission}, confirming the double peaked character of the NEL. Finally, two objects, J121342+4222 and J023658+0242 have been observed twice by LAMOST and the data as well as the fitting results of the two-epoch spectra are consistent. Fig \ref{fig4-dis}c shows the distribution of the intensity ratios of the ''blue'' and ''red'' NEL velocity components in \oiii $\lambda 5007$ line. About 50\% objects in our sample have a flux ratio between 0.75 and 1.25. Smith et al. (\cite{smith2012}) argue that double peaked AGN in which the two narrow line components have closely similar intensity often represent rotating disks, and are inconsistent with a black hole binary scenario. There remain 50\% objects, in which the ''blue'' component flux is stronger than the ''red'' component one. If a simple interaction between the a radio jet and the NLR was at work, (see also 3.4.2 below) the ''blue'' component should be systematically stronger than the ''red'' component. Fig \ref{fig4-dis}d shows the distributions of the FWHM of the ''blue'' and ''red'' NEL velocity components. Most objects exhibit intrinsically narrow lines, with FWHM $\leq$ 300 \kms on both components, Fig. \ref{fig5-bpt} shows, for the objects belonging to the {\it galaxy} subsample, the location of the two NEL components in the Baldwin, Philips \& Terlevich (\cite{ref70}) (hereafter BPT) emission line diagnostic diagram. We recall that there is still some debate about the accuracy of the AGN versus starburst (or HII-like) spectrum distinction in the BPT diagram, the theoretical dividing line as derived by Kewley et al. (\cite{ref56}) being largely offset from the empirical dividing line derived by Kauffmann et al.(\cite{ref55}) from a very large sample of emission-line galaxies. In 7 objects, both components are clearly located in the Kewley et al. AGN region, implying that the ionization of their narrow-line plasma are dominated by non-thermal sources. Note that, from this diagram only, radiative shock contribution cannot be evaluated quantitatively. 3 other objects have their two components close to the Kewley et al separation line between AGN-like spectra and thermal HII-like spectra, but inside the Kauffmann et al. AGN region. For these objects, a significant contribution of thermal photoionization by ongoing massive star formation is possible. One object (J083425+283451) exhibits widely different behaviour in its two NEL components, the ''blue'' one being probably largely dominated by thermal photoionization, the ''red'' one being transitional, or AGN-like, on the Kewley et al. separation line. The 4 remaining objects have their NEL components clustered in a small area of the diagram, with a small dispersion around \oiii / \hb = 1 for \nii / \ha = 0.4 , i.e. close to the Kauffmann et al. AGN / HII separation line. These objects are likely to be largely dominated by massive star formation, non-thermal AGN-like ionization playing only a marginal role, if any, but weak shocks being still possible. For all objects and all components, the ionization parameter is quite strong. These results are summarized in Table \ref{tab2-emission}, column 10. \begin{landscape} \begin{table}[htb!] \caption{Emission lines properties of the double-peaked emission line sample} \label{tab2-emission} \centering \setlength{\tabcolsep}{2pt} \renewcommand{\arraystretch}{1.5} \begin{tabular}{@{}ccccccccccc@{}} \hline\noalign{\smallskip} Designation & NEL model & $\bigtriangleup V$ & $FWHM_{b}$ & $FWHM_{r}$ & $F{^b_{\texttt{\oiii}}}$/$F{^b_{H_{\beta}}}$ & $F{^r_{\texttt{\oiii}}}$/$F{^r_{H_{\beta}}}$ & $F{^b_{\texttt{\nii}}}$/$F{^b_{H_{\alpha}}}$ & $F{^r_{\texttt{\nii}}}$/$F{^r_{H_{\alpha}}}$ & Type & Ref\\ \scriptsize(1) & \scriptsize(2) & \scriptsize(3) & \scriptsize(4) & \scriptsize(5) & \scriptsize(6) & \scriptsize(7) & \scriptsize(8) & \scriptsize(9) & \scriptsize(10) & \scriptsize(11)\\ \hline\noalign{\smallskip} J005407.03+073611.9 & 2G &$ 279 \pm 44 $&$ 301 \pm 43 $&$ 224 \pm 48 $&$ 1.42 \pm 0.25 $&$ 1.17 \pm 0.34 $&$ 0.39 \pm 0.03 $&$ 0.40 \pm 0.04 $& SF + SF & - \\ J023658.06+024217.9 & 2G + W &$ 250 \pm 49 $&$ 236 \pm 44 $&$ 241 \pm 33 $&$ 4.99 \pm 0.99 $&$ 5.88 \pm 1.21 $&$ 0.64 \pm 0.04 $&$ 0.75 \pm 0.05 $& AGN + AGN & - \\ J023832.67+023349.1 & 2G &$ 378 \pm 83 $&$ 397 \pm 53 $&$ 325 \pm 39 $&$ 10.20 \pm 1.49 $&$ 12.91 \pm 1.99 $&$ 0.85 \pm 0.11 $&$ 0.72 \pm 0.09 $& AGN + AGN & - \\ J083425.28+283451.3 & 2G &$ 294 \pm 43 $&$ 296 \pm 16 $&$ 346 \pm 53 $&$ 0.37 \pm 0.02 $&$ 2.31 \pm 0.54 $&$ 0.48 \pm 0.01 $&$ 0.66 \pm 0.01 $& SF + Comp & (4b) \\ J091646.03+283526.7 & 2G + W &$ 392 \pm 17 $&$ 391 \pm 15 $&$ 251 \pm 29 $&$ 5.39 \pm 0.14 $&$ 6.05 \pm 0.44 $&$ 0.56 \pm 0.01 $&$ 0.35 \pm 0.02 $& AGN + AGN & (2) \\ J094430.79+435421.4 & 2G + W &$ 306 \pm 16 $&$ 280 \pm 15 $&$ 255 \pm 21 $&$ 4.81 \pm 1.03 $&$ 9.24 \pm 4.32 $&$ - $&$ - $& - & - \\ J100708.01+242039.0 & 2G + W &$ 340 \pm 25 $&$ 373 \pm 80 $&$ 230 \pm 115 $&$ 3.87 \pm 1.74 $&$ 3.43 \pm 0.71 $&$ - $&$ - $& - & (3) \\ J104718.31+254348.3 & 2G + W &$ 249 \pm 58 $&$ 252 \pm 46 $&$ 270 \pm 59 $&$ 6.31 \pm 2.27 $&$ 13.38 \pm 9.54 $&$ - $&$ - $& - & - \\ J113630.61+135848.8 & 2G &$ 192 \pm 39 $&$ 322 \pm 33 $&$ 301 \pm 30 $&$ 2.70 \pm 0.39 $&$ 1.94 \pm 0.18 $&$ 0.77 \pm 0.02 $&$ 0.79 \pm 0.01 $& Comp + Comp & (1) \\ J121342.90+422202.8 & 2G &$ 349 \pm 88 $&$ 414 \pm 72 $&$ 425 \pm 44 $&$ 9.30 \pm 0.54 $&$ 7.36 \pm 0.43 $&$ 1.20 \pm 0.01 $&$ 1.06 \pm 0.02 $& AGN + AGN & - \\ J123314.49+262624.9 & 2G + W &$ 253 \pm 16 $&$ 297 \pm 23 $&$ 239 \pm 17 $&$ 7.56 \pm 0.66 $&$ 7.41 \pm 0.63 $&$ - $&$ - $& - & - \\ J131434.73+563419.3 & 2G &$ 181 \pm 21 $&$ 282 \pm 25 $&$ 251 \pm 19 $&$ 0.89 \pm 0.05 $&$ 0.91 \pm 0.06 $&$ 0.43 \pm 0.01 $&$ 0.37 \pm 0.01 $& SF + SF & (4b) \\ J133730.29-002525.4 & 2G &$ 239 \pm 18 $&$ 267 \pm 19 $&$ 230 \pm 18 $&$ 2.39 \pm 0.21 $&$ 3.33 \pm 0.31 $&$ 0.64 \pm 0.02 $&$ 0.46 \pm 0.02 $& Comp + Comp & (4b) \\ J133737.82+555816.7 & 2G + W &$ 265 \pm 28 $&$ 314 \pm 21 $&$ 327 \pm 31 $&$ 8.92 \pm 4.83 $&$ 3.05 \pm 0.65 $&$ - $&$ - $& - & - \\ J135207.73+052555.8 & 2G &$ 331 \pm 16 $&$ 450 \pm 20 $&$ 263 \pm 9 $&$ 3.74 \pm 0.15 $&$ 7.36 \pm 0.36 $&$ 1.52 \pm 0.03 $&$ 1.83 \pm 0.03 $& AGN + AGN & (1), (4a) \\ J135646.10+102609.0 & 2G + W &$ 386 \pm 23 $&$ 484 \pm 23 $&$ 394 \pm 26 $&$ 6.59 \pm 0.46 $&$ 8.29 \pm 0.57 $&$ 0.59 \pm 0.02* $&$ 0.34 \pm 0.01* $& AGN + AGN & (2) \\ J140225.68+465817.4 & 2G &$ 224 \pm 42 $&$ 281 \pm 20 $&$ 260 \pm 24 $&$ 1.03 \pm 0.04 $&$ 0.51 \pm 0.02 $&$ 0.38 \pm 0.01 $&$ 0.35 \pm 0.01 $& SF + SF & - \\ J140231.28+184807.8 & 2G + W & $ 156 \pm 109 $&$ 203 \pm 74 $&$ 207 \pm 72 $&$ 3.86 \pm 1.58 $&$ 3.39 \pm 1.09 $&$ 0.44 \pm 0.08 $&$ 0.35 \pm 0.05 $& Comp + Comp & - \\ J150501.56+371311.7 & 2G &$ 256 \pm 39 $&$ 287 \pm 19 $&$ 292 \pm 37 $&$ 4.81 \pm 0.23 $&$ 9.33 \pm 1.25 $&$ 0.58 \pm 0.01 $&$ 0.73 \pm 0.02 $& AGN + AGN & (4b) \\ J232703.17+004256.7 & 2G &$ 267 \pm 18 $&$ 332 \pm 16 $&$ 235 \pm 16 $&$ 1.11 \pm 0.05 $&$ 1.16 \pm 0.06 $&$ 0.36 \pm 0.01 $&$ 0.38 \pm 0.01 $& SF + SF & (4a) \\ \hline\hline \end{tabular} \tablecomments{0.86\textwidth}{ Column 1: LAMOST DR1 designation hhmmss.ss+ddmmss.s (J2000.0)\\ Column 2: best NEL model: 2G: 2 gaussians, 2G+W : 2 gaussian + a broader gaussian ''wing''\\ Column 3: velocity difference in \kms between the two NEL components, in the rest frame defined by the redshift listed in Table \ref{tab1-list} \\ Columns 4-5: FWHMs of blue and red components, in units of \kms \\ Column 6-7: flux ratio of \oiii\ $\lambda\ 5007$ and \hb \\ Column 8-9: flux ratio of \nii\ $\lambda\ 6583$ and \ha. Note * indicate that the values are from the spectrum of SDSS for technical reason. \\ Column 10: NEL components location in BPT diagram: AGN: AGN-like line ratios, SF: likely dominated by massive star formation, Comp: on the Kewley et al. (\cite{ref56}) AGN-HII separation line, possibly composite or transitional (in the Kauffmann et al. (\cite{ref55}) these would be classified AGN) \\ Column 11: (1): Wang et al. \cite{ref5}, (2): Liu et al. \cite{ref65}, (3): Smith et al. \cite{ref8}, (4) Ge et al. \cite{ref9}: a) listed as double-peaked NEL, b) listed as asymmetric NEL } \end{table} \end{landscape} \begin{table}[htb!] \caption[]{Additional FIRST radio data for our double-peaked NEL sample from LAMOST DR1} \label{tab3-first} \centering \begin{tabular}{cccccc} \hline \hline\noalign{\smallskip} Designation & Max. flux density & Integrated flux & s/n & Note & Ref. \\ \scriptsize(1) & \scriptsize(2) & \scriptsize(3) & \scriptsize(4) & \scriptsize(5) & \scriptsize(6)\\ \hline\noalign{\smallskip} J005407.03+073611.9 & 0.46 & 0.48 & 4.0 & P & (2) \\ J023658.06+024217.9 & 0.36 & 0.38 & 3.0 & P & (2) \\ J023832.67+023349.1 & 23.2 & 24.94 & 142 & Ext. & (1),(3) \\ J091646.03+283526.7 & 5.25 & 5.76 & 35 & P & (1),(3) \\ J094430.79+435421.4 & 0.61 & 0.94 & 5.0 & Ext?, (*) & (2) \\ J113630.61+135848.8 & 0.42 & 0.65 & 3.5 & P & (2) \\ J121342.90+422202.8 & 1.84 & 1.58 & 12 & P & (1) \\ J131434.73+563419.3 & 0.74 & 0.99 & 5.2 & P & (2) \\ J135207.73+052555.8 & 0.64 & 1.26 & 4.5 & P & (2) \\ J135646.10+102609.0 & 56.5 & 59.58 & 362 & P & (1),(3) \\ J140225.68+465817.4 & 0.64 & 0.52 & 4.5 & P & (2) \\ J140231.28+184807.8 & 0.90 & 1.04 & 6.2 & P & (1) \\ \noalign{\smallskip} \hline \end{tabular} \tablecomments{0.86\textwidth}{ Column 1: LAMOST DR1 designation hhmmss.ss+ddmmss.s (J2000.0); Column 2: Maximum flux density on source, in $mJy/beam$; Columns 3: integrated flux in $mJy$ from FIRST catalog; Column 4: local signal-to-noise on source; Column 5: P: point-source, Ext: resolved extended source; Column 6: reference: (1): FIRST source catalog, (2): present work, (3) also in NVSS \\ (*) indicates possible sidelobe contamination on FIRST map. } \end{table} \begin{figure} \centering \subfigure{ \begin{minipage}[b]{0.48\textwidth} \includegraphics[width=1\textwidth]{ms1711fig4a.eps} \\ \includegraphics[width=1\textwidth]{ms1711fig4c.eps} \end{minipage} } \subfigure{ \begin{minipage}[b]{0.48\textwidth} \includegraphics[width=1\textwidth]{ms1711fig4b.eps} \\ \includegraphics[width=1\textwidth]{ms1711fig4d.eps} \end{minipage} } \caption{The distribution of our double-peaked objects. Top left (a): redshift distribution of galaxies(empty bars) and QSOs(line bars); Top right (b): distribution of velocity difference between NEL components; Bottom left (c): flux ratio of the red and blue \oiii\ components; Bottom right (d): FWHM distribution of blue (empty bars) and red (line bars) components;} \label{fig4-dis} \end{figure} \subsection{Additional radio data at 1.4 $GHz$ and WISE infrared data} We have searched the FIRST radio survey for counterparts at 1.4 $GHz$ of our candidates. 5 sources have been found in the FIRST source catalog, whose flux limit is on average 1 $mJy$ across the surveyed area. However, from a close examination of the FIRST image cutouts, 7 more detections were found, all with close positional coincidences (less than 2 ''). 2 of these remain quite marginal and the derived fluxes may have large uncertainties. All sources are unresolved except J023832+023349 (alias PKS 0235+023) which is a radiogalaxy exhibiting structure in its image. J094430+435421 has a slightly NS elongated radio image but appears located in a quite noisy area polluted by ripples that could be due to sidelobes effects from a distant bright source in the same field. The radio data derived from FIRST are listed in Table \ref{tab3-first}. We have also examined the NVSS survey, but the angular resolution is much better in FIRST (5 '' instead of 45 ''), although the sensitivity of NVSS to weak extended structures may be better. In the NVSS catalog, 3 sources are present, namely J023832+023349, J091646+283526 and J135646+102609.\\ \indent The WISE space experiment (Wright et al. \cite{ref67}, Mainzer et al. \cite{ref68}) has provided an all-sky map and source catalog in 4 spectral bands centered at 3.4, 4.6, 12 and 22 microns. The two first bands are usually dominated by starlight, the two longest by hot interstellar dust. The spatial resolution is excellent ( ~6" in the three shortest bands, 12" in the fourth one) and the sensitivity very high. We have searched the AllWISE catalog for counterparts to our objects. The results are given in Table \ref{tab4-wise}, where WISE magnitudes have been converted to fluxes in millijanskys using the recommendations of Jarrett et al. (\cite{ref69}). All sources are detected in the first three bands. In the 22 $\mu$ band 4 objects have a s/n below the formal detection threshold. 3 sources (J023658.06+024217.9, J121342.90+422202.8, J135207.73+052555.8) are fitted with two gaussian point-source components by the catalog construction pipeline, implying resolved structure. \begin{table}[htb!] \caption[]{Additional WISE infrared data for our double-peaked NEL sample from LAMOST DR1} \label{tab4-wise} \centering \begin{tabular}{ccccc} \hline \hline\noalign{\smallskip} Designation & 3.4 $\mu$ flux & 4.6 $\mu$ flux & 12 $\mu$ flux & 22 $\mu$ flux \\ \scriptsize(1) & \scriptsize(2) & \scriptsize(3) & \scriptsize(4) & \scriptsize(5) \\ \hline\noalign{\smallskip} J005407.03+073611.9 & $ 0.79 \pm 0.02 $ & $ 0.60 \pm 0.02 $ & $ 3.22 \pm 0.18 $ & $ 8.00 \pm 1.21 $ \\ J023658.06+024217.9 & $ 2.20 \pm 0.07 $ & $ 1.50 \pm 0.06 $ & $ 7.55 \pm 0.35 $ & $ 9.70 \pm 1.80 $ \\ J023832.67+023349.1 & $ 4.00 \pm 0.09 $ & $ 5.49 \pm 0.12 $ & $ 11.84 \pm 0.25 $ & $ 34.15 \pm 1.40 $ \\ J083425.28+283451.3 & $ 0.80 \pm 0.02 $ & $ 0.57 \pm 0.02 $ & $ 2.45 \pm 0.16 $ & undetected \\ J091646.03+283526.7 & $ 5.17 \pm 0.11 $ & $ 7.75 \pm 0.15 $ & $ 23.06 \pm 0.41 $ & $ 68.62 \pm 2.37 $ \\ J094430.79+435421.4 & $ 0.97 \pm 0.02 $ & $ 1.35 \pm 0.03 $ & $ 2.22 \pm 0.15 $ & $ 6.85 \pm 1.16 $ \\ J100708.01+242039.0 & $ 0.69 \pm 0.02 $ & $ 1.00 \pm 0.03 $ & $ 1.88 \pm 0.15 $ & $ 4.25 \pm 1.37 $ \\ J104718.31+254348.3 & $ 0.54 \pm 0.02 $ & $ 0.88 \pm 0.02 $ & $ 1.63 \pm 0.18 $ & $ 6.52 \pm 1.18 $ \\ J113630.61+135848.8 & $ 1.21 \pm 0.03 $ & $ 0.94 \pm 0.03 $ & $ 3.76 \pm 0.20 $ & $ 5.92 \pm 1.51 $ \\ J121342.90+422202.8 & $ 2.75 \pm 0.06 $ & $ 1.97 \pm 0.05 $ & $ 6.50 \pm 0.25 $ & $ 19.65 \pm 1.33 $ \\ J123314.49+262624.9 & $ 0.40 \pm 0.01 $ & $ 0.56 \pm 0.02 $ & $ 1.25 \pm 0.14 $ & undetected \\ J131434.73+563419.3 & $ 0.71 \pm 0.02 $ & $ 0.66 \pm 0.02 $ & $ 5.39 \pm 0.16 $ & $ 9.68 \pm 0.93 $ \\ J133730.29-002525.4 & $ 0.68 \pm 0.02 $ & $ 0.58 \pm 0.02 $ & $ 3.37 \pm 0.16 $ & $ 5.91 \pm 0.86 $ \\ J133737.82+555816.7 & $ 0.22 \pm 0.01 $ & $ 0.28 \pm 0.01 $ & $ 0.93 \pm 0.12 $ & undetected \\ J135207.73+052555.8 & $ 5.25 \pm 0.15 $ & $ 2.53 \pm 0.10 $ & $ 6.63 \pm 0.30 $ & $ 9.50 \pm 1.80 $ \\ J135646.10+102609.0 & $ 2.23 \pm 0.05 $ & $ 4.47 \pm 0.09 $ & $ 29.19 \pm 0.46 $ & $ 178.83 \pm 4.37 $ \\ J140225.68+465817.4 & $ 0.86 \pm 0.02 $ & $ 0.66 \pm 0.02 $ & $ 5.18 \pm 0.16 $ & $ 8.42 \pm 0.77 $ \\ J140231.28+184807.8 & $ 0.36 \pm 0.01 $ & $ 0.48 \pm 0.02 $ & $ 2.45 \pm 0.12 $ & $ 9.16 \pm 0.82 $ \\ J150501.56+371311.7 & $ 3.93 \pm 0.09 $ & $ 4.67 \pm 0.09 $ & $ 11.42 \pm 0.25 $ & $ 27.30 \pm 1.04 $ \\ J232703.17+004256.7 & $ 1.06 \pm 0.03 $ & $ 0.68 \pm 0.02 $ & $ 1.77 \pm 0.17 $ & undetected \\ \noalign{\smallskip} \hline \end{tabular} \tablecomments{0.86\textwidth}{ Column 1: LAMOST DR1 designation hhmmss.ss+ddmmss.s (J2000.0); Column 2-5: WISE flux of source, in $mJy$, computed from the magnitudes in AllWISE Catalog, for passbands 3.4, 4.6, 12 and 22 $\mu$. \\ For three sources resolved in subcomponents, the fluxes of components were added. The 22 $\mu$ fluxes have been corrected following Jarrett et al. (\cite{ref69}). A source is considered as undetected if its s/n is lower than 3 \\ } \end{table} \subsection{Notes on individual objects} We report here some remarks from SDSS images and references on previous work related to two objects. J005407+073611: no morphological information available (technical problems in SDSS image) J023658+024217: an Sb spiral with a bright nucleus. In close group with several galaxies whose appearance suggest similar redshift. The absorption lines could also be double, but this needs confirmation with higher s/n. The SDSS redshift corresponds to the velocity of the blue NEL component. J023832+023349, (PKS 0235+023), is a broad line Seyfert 1 galaxy, probably of SBa type, whose \ha\ broad line exhibits double peaks or twin shoulders. This was fitted quite well, on 1991 observations, with a model attributing the broad emission to a circular, relativistic, Keplerian disk (Eracleous \& Halpern \cite{ref37, ref49}). However, Gezari et al. (\cite{ref50}) presented a long-term monitoring of the double-peaked broad \ha\ emission lines and found a dramatic change in profile shape after 1991 leading to consider the simple circular disk model as insufficient for interpretation of the spectra taken after 1991. The present LAMOST spectrum not only shows a double peak on \ha\ broad component, but also double peaks on all $narrow$ lines including [OII], [NeIII] and [NeV] as well as the Balmer series. The broad component seems highly obscured by internal extinction, since it is almost invisible in \hb. It is clear that this object deserves detailed follow-up observations. J083425+283451: a SBb spiral with a low-excitation emission spectrum. J091646+283526 is a quite amorphous galaxy with a perturbed morphology. The bright nucleus is close (2-3'') to a second, much redder one. A moderately broad component is present in the Balmer emission lines. The excitation is very strong with bright HeII and [NeV] lines (Seyfert 1.5). This object is a very good merger candidate. The NVSS map shows an extended source with two components, one being $\pm$ coincident with the optical object. Fu et al. (\cite{ref22}) place J091646+283526 in the category of unresolved narrow-line regions from integral-field spectroscopy. They argue that most of these spatially unresolved double-peaked NEL AGNs are aligned or young outflows. J094430+435421: QSO whose NELs are well splitted. The \hb\ line is coincident with the telluric A band of oxygen making spectrophotometry uncertain. J100708+242039 and J104718+254348 are QSOs on which few information is available. The latter is an X-ray source. J113630+135848: a $\sim$ amorphous galaxy with possible pair of opposite, low surface brightness, blue diffuse extensions. J121342+422202: a galaxy with a distended asymmetric envelope. The \ha\ range is lacking on the SDSS spectrum. A small diffuse object is very close. The galaxy is member of a group of more than 10 objects of similar redshift, among which a Seyfert 1 at 4.1 $arcmin$ N. Possible merger or post-merger. J123314+262624: QSO, no information. J131434+563419: peculiar asymmetric-shaped, possibly barred, galaxy with a blue extension that could be a tidal tail. Possible merger candidate. In a loose group with other, similar redshift, galaxies. J133730-002525: a spiralgalaxy (Sbc/Sc) seen close to face-on, with many faint objects close-by (but no redshift information on them, except one at 5 $arcmin$ NE) J133737+555816: QSO, no information J135207+052555: typical luminous Seyfert 2 galaxy, of early-type (SO or SO/a). The nucleus appears unique at the spatial resolution of SDSS image. The double-lined character of the NEL is especially obvious on the \sii\ lines which appear as a triple peak. This object is a member of the Abell 1809 galaxy cluster. J135646+102609 has aroused great interest in the recent literature. Liu et al. (\cite{ref6}) included this object in a sample of type 2 AGNs with double-peaked NEL. Its morphology is clearly disturbed, suggesting an ongoing merger. In a loose cluster with several other less luminous objects. The red NEL component has the same redshift as the absorption lines of the stellar population. The presence of two merging galactic nuclei has been demonstrated from optical (Greene et al. \cite{ref51, ref52}; Fu et al. \cite{ref22}) and near-infrared (Shen et al. \cite{ref24}) observations. Greene et al. (\cite{ref53}) have published two long-slit Magellan LDSS-3 spectra and additional spectral band data. They consider that this object is a pair of interacting galaxies that hosts a luminous obscured quasar in its northern nucleus. Note that the optical spectrum from LAMOST or SDSS is that of a typical unobscured Seyfert 2 galaxy, all lines being narrow. J140225+465817: a galaxy with perturbed morphology, with an extension (or projected close companion) along NE and irregular fuzzy extension along S. Possible merger candidate. Isolated. J140231+184807: small apparent diameter galaxy, seems isolated. J150501+371311: bright nucleus galaxy with a pear-shaped disk. A diffuse blue object at 25 $arcsec$ E, without redshift information. J232703+004256: an $\sim$ amorphous galaxy (maybe Sa), forming an interactive pair with a companion at 25 $arcsec$ NE, of similar size and brightness. A diffuse matter bridge is seen between both objects, and a tidal tail appears towards SW. Single nucleus at SDSS resolution. \subsection{Possible origins for double NEL structures} \subsubsection{Black hole binaries} One initial aim of this investigation was the quest for QSO and AGN candidates hosting binary black holes from the identification of double-peaked NEL in optical spectra, as practised by Wang et al. (\cite{ref5}), Liu et al. (\cite{ref6,ref7}) and Smith et al. (\cite{ref8}). In our sample appeared galaxy nuclei candidates with consistent ''blue'' and ''red'' velocity systems in all the emission lines (i. e., \hb, \oiii\ $\lambda\lambda\ 4959, 5007$, \nii $\lambda\lambda\ 6548, 6584$, \ha\ and \sii $\lambda\lambda\ 6717, 6731$ ) of which five examples are shown in Fig 2. Among the 15 galaxies identified, 6 exhibit pairs of NEL components that may be unambiguously classified as AGN. The question remains if these double NEL systems correspond to double supermassive black hole engines. Besides, 5 QSO are also identified, among which one (J094430+435421) has spectacular double NEL components and has only a faint radio loudness. On a general ground, statistical analysis of double-peaked \oiii\ AGN samples by Shen et al. (\cite{ref24}) and Fu et al. (\cite{ref26}) tend to support the conclusion that a large fraction of the double-peak systems are driven by the NLR dynamics instead of the existence of black hole binaries. Ge et al. (\cite{ref9}) found only 54 dual-core galaxies with projected separations closer than 3" among a sample of 15600 objects with double-peak or asymmetric NEL, but the angular resolution of SDSS images is not appropriate to disentangle very close components. We also stress that previous studies of double-peaked NEL have focused on very obvious, well separated double lines in the line profiles, which have been observed since long but remain largely unexplained. In the present limited sample, with the availablable low or moderate spatial and spectral resolutions (optical spectra, SDSS images and FIRST radio maps), we do not find evidence for binary nuclei as exhibited by the already well-known binary BH candidates (Ge et al. \cite{ref9}), except in J091646+2835, which has two central condensations and has a double AGN-like spectrum NEL component. (Note that the well-studied J135646+102609 does not show up clearly as a binary nucleus merger on the SDSS images). This study hence must be completed with additional higher resolution including good seeing imagery and other multi-band data from follow-up observations. \subsubsection{Interaction between jet and NLR} In radio-loud AGNs, double narrow line structure may readily arise from the radio-jet interaction with the NLR clouds. This operates mainly by ram pressure acceleration of the ionized gas clouds of the NLR, and can produce complicated NLR kinematics, and especially double lines, due to projection on the line of sight of cone structure in the matter submitted to the interaction. If the jet (whose direction points $\pm$ closely towards the observer) drags and accelerates NLR matter with it, a Doppler blueshifted (with respect to the systemic redshift) ionized gas component may naturally appear in the spectrum. Note that Smith et al. (\cite{ref8}) exclude the radio-loud AGNs from their candidates binary black hole because they think that these objects are more or less dominated by the jet / NLR interaction. The literature on these phenomena is quite extensive, and there have been extremely detailed studies on some nearby cases, most of them are Seyfert 2 galaxies (See e.g. Whittle et al. \cite{ref42}, Rosario et al. \cite{ref43} with many references inside), for which high spatial resolution optical and radio mapping is possible. In our double peaked NEL sample, the objects detected at 1.4 $GHz$ that have the largest radio flux (corresponding more or less to the largest radio luminosity) have their NEL components of type AGN + AGN with FWHM values larger than the average (typically larger than 350 \kms ). It should be useful to explore if this trend still holds on larger samples. \subsubsection{Interaction with companion galaxies and merger candidates} Galaxy collisions and subsequent merging are natural sources of double emission lines phenomena, depending on the spatial integration of the spectrograph entrance aperture, on the angle of projection of the interacting system on the line of sight and on the collision phase observed. Since interactions between gas-rich systems are likely to enhance the star formation, ionized gas could be observed in the two components of a strongly interacting system with velocity separations up to a few hundreds \kms. These interactions are accompanied by many morphological perturbations, such as tidal tails, matter bridges, deformation of disks, envelope asymmetric extensions, etc... Also, several authors have proposed that nuclear activity may sometimes be triggered by galaxy interactions. In the present sample, examination of the SDSS images lead to find one clear interacting pair (J232703+0042), 1 merger (J135646+1026), 4 candidates mergers or possible candidates mergers (J091646+2835, J121342+4222, J131434+5634, J140225+4658), 2 more objects have perturbed morphology (J113630+1358, J150501+3713) and J135207+0525 belongs to a populated cluster in which encounters should have higher probability than in the field. Thus, 2/3 of our $galaxy$ sample are objects for which interaction is a plausible origin for NEL double structure. Verification of this hypothesis, its extension to other double-peaked NEL galaxy samples and better modeling cannot be done without follow-up observations, including deep imagery. Regarding the QSO, for which morphology information is not available, several previous authors (Hutchings et al. \cite{ref44}; Malkan, \cite{ref45}; Green and Yee \cite{ref46}) have noted that low-redshift quasars very often have faint companion galaxies with small projected angular separations. A scenario involving a collision between a giant AGN host galaxy and a gas-rich star-forming dwarf may lead to double structure in the NEL of the AGN: in the process of interaction, the companion is disrupted by the tidal field of the giant galaxy. Clumps of its interstellar matter may approach the AGN sufficiently close such that these clumps may be submitted to violent differential accelerations able to increase their velocity dispersion up to significantly higher values than the usual average quiet disk value. If interaction with a neighbor were the origin of the double structure in an AGN NEL, observations could disentangle the ionized gas of the companion, expected to have line ratios consistent with a normal massive stars photoionization, from the NLR component of the AGN in which the ionization is basically due to non-thermal source spectra. The present sample does not enable this for the QSO, because \ha\ is not observed. \begin{figure} \centering \includegraphics[width=0.8\textwidth]{ms1711fig5.eps} \caption{The BPT diagnostic diagram for objects with measured \nii\ and \ha\ emission lines (14 objects). For each object, a blue triangle indicates the low velocity component and the red square indicates the high velocity one. Two components belonging to the same source are connected by a thin continuous black line. The error bars are shown in magenta. The solid curve defined by Kewley et al. (\cite{ref56}) and the dashed curve defined by Kauffmann et al. (\cite{ref55}) show the separation between star-forming galaxies, composite galaxies, and AGNs. The horizontal dotted line are defined by \oiii\ / \hb\ = 3. This line is often suggested to separate Seyferts from LINERS.} \label{fig5-bpt} \end{figure} | 14 | 4 | 1404.7218 |
1404 | 1404.7168_arXiv.txt | Comet C/2013 A1 (Siding Spring) will have a close encounter with Mars on October 19, 2014. We model the dynamical evolution of dust grains from the time of their ejection from the comet nucleus to the Mars close encounter, and determine the flux at Mars. Constraints on the ejection velocity from Hubble Space Telescope observations indicate that the bulk of the grains will likely miss Mars, although it is possible that a few-percent of grains with higher velocities will reach Mars, peaking approximately 90--100 minutes after the close approach of the nucleus, and consisting mostly of millimeter-radius grains ejected from the comet nucleus at a heliocentric distance of approximately 9~AU or larger. At higher velocities, younger grains from sub-millimeter to several millimeter can reach Mars too, although an even smaller fraction of grains is expected have these velocities, with negligible effect on the peak timing. Using NEOWISE observations of the comet, we can estimate that the maximum fluence will be of the order of $10^{-7}$ grains/m$^2$. We include a detailed analysis of how the expected fluence depends on the grain density, ejection velocity, and size-frequency distribution, to account for current model uncertainties and in preparation of possible refined model values in the near future. | Comet C/2013 A1 (Siding Spring) will have a close approach (hereafter c/a) with Mars on October 19, 2014 at approximately 18:29 UT, reaching a minimum distance of approximately 134,000~km from the center of Mars, according to JPL solution \#46 \citep{Farnocchia_in_prep}. This orbital solution was retrieved from the JPL Horizons system \citep{1996DPS....28.2504G}, on March 17, 2014, and include all the observations before the low solar elongation period. The comet is on a hyperbolic retrograde orbit (129$^\circ$ inclination), and the encounter with Mars will be at a relative velocity of approximately 56 km/s. Observed from Mars, the comet reaches a minimum solar elongation of approximately 72$^\circ$. After the close approach between the two bodies, Mars continues to move closer to the orbit of the comet, crossing the comet orbit plane approximately 100 minutes after c/a and then reaching a minimum distance of approximately 27,000~km from the comet orbit approximately 102 minutes after c/a. A detailed trajectory analysis by \cite{Farnocchia_in_prep} indicates that the current uncertainty on the c/a distance, projected on the target plane normal to the velocity of the comet nucleus relative to Mars, is an ellipse approximately $1600 \times 700$~km (1$\sigma$) with the long axis 20$^\circ$ from the line connecting the comet nucleus and Mars, and the c/a time uncertainty is about 1 minute (1$\sigma$). These uncertainties exclude any collision of the comet nucleus with Mars, yet warrant a detailed investigation of the possibility that the dust coma of the comet may be able to reach Mars. C/2013~A1 was discovered at a heliocentric distance of 7.2~AU \citep{2013CBET.3368....1M}, suggesting hyper-volatile activity (i.e., driven by CO or CO$_2$) was likely causing long-term build-up of larger grains in the coma. Hubble Space Telescope (HST) observations by \cite{Li_in_prep} allow us to estimate the ejection velocity of the dust grains at the $\mu$m radius scale, or more precisely grains with $\beta\simeq1$, where $\beta$ is the ratio between the radiation pressure force and the gravitational force due to the Sun on the grain \citep{1979Icar...40....1B}. Observations were performed on October 29, 2013, on Jan 21, 2014, and on March 11, 2014. By using a technique based on the sunward turn back distance in the continuum \citep[e.g.,][]{2013Icar..222..799M} we determined a velocity approximately between 28 and 46~m/s at a heliocentric distance of 4.59~AU, then between 44 and 74~m/s at 3.77~AU, and between 48 and 80~m/s at 3.28~AU, all for $\beta=1$ dust grains, see Figure~\ref{fig:v_ej_micron_vs_Rh}. The ranges are relatively broad because of the inherent difficulty of determining a single value for the turn back distance in a continuous brightness distribution. \begin{figure}[t] \begin{center} \includegraphics*[width=0.7\textwidth]{vref.eps} \end{center} \caption{Ejection velocity of dust grains with $\beta=1$ as a function of the heliocentric distance of the nucleus at the time of the ejection. The curves are not a fit, but correspond to the empirical scaling law of Eq.~\eqref{eq:v_ej} with $v_\text{ref}$ values between 0.5~m/s and 1.5~m/s in increments of 0.1~m/s. The three vertical bars are the velocity ranges as constrained from HST observations \citep{Li_in_prep}. } \label{fig:v_ej_micron_vs_Rh} \end{figure} NEOWISE obtained infrared observations of C/2013~A1 on Jan 16, 2014 \citep{Mainzer_in_prep}, when the comet was at 3.82~AU from the Sun. We used these observations to model the dust production. Adopting the techniques in \cite{2011ApJ...738..171B} yields a dust production rate of approximately 10~kg/s at the grain velocities for that heliocentric distance, and for grains with radius between 1.7~$\mu$m and 2.3~$\mu$m \citep{Mainzer_in_prep}. In this paper we use the HST and NEOWISE observations to anchor the grain ejection velocity and production rate. In the next Section we discuss how we use these constraints to model the velocity of larger dust grains ejected at larger heliocentric distances, which are in the relevant range for delivery of dust grains to Mars. In Section~\ref{SEC:RESULTS}, we present the results from numerical simulations where we model the dynamical evolution of dust grains after ejection from the comet nucleus to their close approach with Mars. In Section~\ref{SEC:DISCUSSION} we discuss the consequences of our modeling choices and how they likely affect the outcomes of the simulations, and compare our results with earlier works which characterize the delivery of dust grains to Mars \citep{2014Icar..231...13M,2014MNRAS.439.3294V,2014arXiv1403.7128Y}. Conclusion are then drawn in Section~\ref{SEC:CONCLUSIONS}. | \label{SEC:CONCLUSIONS} We have presented the results of our dynamical model for the delivery of dust grains ejected from the nucleus of comet C/2013~A1 (Siding Spring) to Mars. Direct extrapolation of the grain ejection velocity from HST observations indicates that the bulk of the grains will likely miss Mars. If we include the possibility of a few-percent of the grains to have higher velocities, we find that millimeter radius grains ejected at 9~AU or more from the Sun will collide with Mars. At higher velocities, younger grains from sub-millimeter to several millimeter can reach Mars too. The maximum fluence will be of the order of $10^{-7}$ grains/m$^2$. The timing of the peak flux is expected to be 90--100 minutes after c/a, which is between 19:59 and 20:09 UT of October 19, 2014. | 14 | 4 | 1404.7168 |
1404 | 1404.2051_arXiv.txt | {} {Following the recent detection of {\tArhp} in the Crab nebula spectrum % , we have computed the photodissociation rate of {\Arhp} in order to constrain the physical processes at work in this environment.} {Photodissociation cross sections of {\Arhp} are computed in an {\ai} approach including explicit account of spin-orbit coupling.} { We report the photodissociation cross section of {\Arhp} as a function of wavelength. Photodissociation probabilities are derived for different impinging radiation fields.% The photodissociation probability of for a very small unshielded cloud surrounded on all sides by the unshielded InterStellar Radiation Field (ISRF) model described by \cite{draine:78} is equal to $9.9 \cdot10^{-12}$ s$^{-1}$ and 1.9 $\cdot 10^{-9}$ s$^{-1}$ in the Crab nebula conditions. The dependence on the visual extinction is obtained by using the Meudon Photon Dominated Region (PDR) code and corresponding analytical fits are provided.} { These data will help to produce a realistic chemical network to interpret the observations. Photodissociation of {\Arhp} is found to be moderate and the presence of this molecular ion is mainly dependent on the molecular fraction.} | Whereas {\Arhp} has been studied in great detail from laboratory and theoretical techniques, the detection of the {\tArhp} isotopologue in the Crab nebula \citep{barlow:13} offers an outstanding opportunity to exploit these studies in a previously unforeseen astrophysical context. The main motivation for this article is to investigate the response of this molecular ion to the high ambient radiation field and subsequently check the chemical processes at work in the Crab Nebula environment. Section~\ref{sec:theory} describes the recent {\ai} studies of {\Arhp}. We use the available information to compute photodissociation rates resulting from different shapes of the incident radiation fields as discussed in Section~\ref{sec:pdr}. We first describe the main features of Photon Dominated Regions (PDR) models and introduce different radiation fields to compute the photodissociation probabilities. We summarize our conclusions in Section~\ref{sec:conclusion}. | \label{sec:conclusion} We report photodissociation cross sections of {\Arhp} as a function of wavelength which can be used to compute the response of that molecular ion to ultraviolet interstellar radiation fields. The photodissociation rate of {\Arhp} for a very small unshielded cloud surrounded on all sides by the standard Draine UV ISRF is $9.9\cdot10^{-12}$ s$^{-1}$. This value is very sensitive to the spectral distribution % in the VUV 911 - 2400 $\AA$ wavelength window, as shown in Table \ref{tab:res}. % The corresponding rate relevant to the Crab nebula environment % is 1.9 $\cdot$ 10$^{-9}$ s$^{-1}$. The dependence on the visual extinction A$_V$ due to the attenuation of radiation by dust is derived from the Meudon PDR code both for incident isotropic and plane parallel radiations. We report the corresponding analytic functions for a semi infinite plane parallel cloud for different typical cases and for the Crab nebula environment. The constant numerical factor is close to half the value obtained for the unshielded environment. The dependence on A$_V$ is better represented with an E$_2$ function when isotropic radiation is impinging on the cloud. Nevertheless, we stress out that a direct integration of the product of photodissociation cross sections by the actual radiation field allows to derive photo destruction rates unambiguously. The actual value of the photodissociation rate is moderate and destruction of {\Arhp} is mainly due to molecular \HH. Then, {\Arhp}, as well as {\Ohp} also detected in the Crab nebula \citep{barlow:13}, are strongly dependent on the molecular fraction of the gas. % | 14 | 4 | 1404.2051 |
1404 | 1404.6672_arXiv.txt | Inspired by quantum gravitational physics, the approach of non-commutative (NC) phase space leads to a modified dispersion relation of gravitational waves. This feature, if applied to the very early universe, gives rise to a modified power spectrum of primordial tensor perturbations with a suppression of power on large scales. We confront this phenomenon with the BICEP2 and Planck experiments, and show that inflation with the modified dispersion relation can simultaneously fit the observations better than the standard inflationary paradigm. In particular, the numerical result implies that with the latest cosmological microwave background (CMB) observations, a quantum gravity modified power spectrum of primordial tensor modes is preferred at a statistical significance of more than $3\sigma$ compared with the minimal model. Our study indicates that the potential tension between the BICEP2 and Planck data may be resolved by quantum gravity effects. | 14 | 4 | 1404.6672 |
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1404 | 1404.4368.txt | The \starname system consists of five planets orbiting an early M dwarf. The planets have physical radii of 1.0--1.50 \rearth and orbital periods of 4--130 days. The $1.1 \rearth$ \planetf with a period of 130\,days is of particular interest. Its insolation of roughly $0.32$ \insol places it within the surface liquid water habitable zone (HZ). We present a multifaceted study of the \starname system, using two sets of parameters which are consistent with the data and also self-consistent. First, we show that the distribution of planet masses can be roughly reproduced if the planets were accreted from a high surface density disk presumably sculpted by an earlier phase of migration. However, our simulations predict the existence of one to two undetected planets between planets e and f. Next, we present a dynamical analysis of the system including the effect of tides. The timescale for tidal evolution is short enough that the four inner planets must have small obliquities and near-synchronous rotation rates. The tidal evolution of \planetf is slow enough that its current spin state depends on a combination of its initial spin state, its dissipation rate, and the stellar age. Finally, we study the habitability of \planetf with a one-dimensional climate model. The planet's surface temperature can be raised above 273 K with 0.5--5 bars of $\mathrm{CO_2}$, depending on the amount of $\mathrm{N_2}$ present. \planetf represents a case study of an Earth-sized planet in the cooler regions of the HZ of a cool star. | The \kepler mission~\citep{borucki2010} has made key discoveries on the road to finding Earth-like planets \citep[e.g.,][]{Batalha.etal.:2011,Borucki.etal:2012,Fressin.etal.2012,Borucki.etal:2013}. The recent detection of an Earth-sized planet in the HZ of an M star (i.e., the \starname system, \citealp{Quintana2014}) brings us a step closer to finding a true Earth twin. \begin{figure}[htbp] \includegraphics[width=0.45\textwidth]{Kepler186_orbits_Seff_k62-eps-converted-to.pdf} \caption{Orbital configuration of the \starname system. The top part shows a top-down view of the system, assuming orbits from set $\mathcal{A}$. The habitable zone boundaries are from \cite{kopparapu2013}: the inner boundaries are the moist/runaway greenhouse limits and the outer boundaries are the maximum greenhouse and early Mars limits. The sizes of the symbols are not to scale with the planetary orbits. The bottom part of the plot shows a comparison between four different planetary systems which contain planets in the HZ: the solar system, Kepler-62~\citep{Borucki.etal:2013}, Kepler-186~\citep{Quintana2014}, and GJ 581~\citep{udry07,mayor09}. Note that the inner moist and runaway greenhouse limits of the habitable zone are the same for Kepler-62, Kepler-186, and GJ~581. Given the consistent insolation scaling, the x axis is linear in orbital distance but the scale is different for each system. The planets' relative sizes are correct, although for GJ~581 the planetary radii were calculated as $R = [M sin(i)]^{2.06}$, following \cite{Lissauer.etal:2011}.} \label{Kepler186_HZ} \end{figure} The \starname planetary system hosts five known planets including \planetf, an Earth-sized planet in the HZ \citep{Selsis2007,kopparapu2013}. Figure \ref{Kepler186_HZ} shows a comparison between the \starname system, the solar system, and two other systems with potentially habitable planets: Kepler-62~\citep{Borucki.etal:2013} and GJ~581~\citep{udry07,mayor09}. Climate models have shown that GJ~581d, a super-Earth near the outer edge of the HZ of its host M star, could sustain surface liquid water \citep[e.g.,][]{wordsworth2010}. \planetf receives a comparable or perhaps slightly higher stellar flux than GJ~581, placing it more comfortably within the HZ. Here, we use the definition of the classical HZ \citep[HZ, e.g., ][]{dole1964,hart1978,kasting1993,Selsis2007,kopparapu2013}. Acknowledging the fact that all terrestrial life needs liquid water, the HZ is defined as the region around a star where a terrestrial planet can host liquid water on the surface. The extent of this HZ naturally depends on the atmospheric conditions (composition, pressure) as well as on the properties of the central star. Many more factors influence the width of the HZ, such as the geological activity \citep[e.g.,][]{lammer2010}, the biosphere itself \citep[e.g.,][]{grenfell2010}, or the dynamical environment of the planetary system \citep[e.g.,][]{menou03,barnes04,jones2006,sandor2007,kopparapu2010}. \medskip We present a three-pronged study of the \starname system. We first try to reproduce the orbital architecture of the system using simple accretion simulations (Section \ref{formation}). We show that certain features of the system--such as the large gap between planets e and f--are hard to explain. We next briefly discuss the long-term dynamical stability of the system (Section \ref{stability}). In Section \ref{tidal}, we study the long-term dynamical, tidal, and spin evolution of the system. We use both simple tidal models and $N$-body simulations which include both tides and general relativity. Next, we study the atmospheric conditions needed to bring \planetf's surface temperature into the liquid water range (Section \ref{habi}). We discuss our findings and conclude in Section \ref{conclusion}. %%%%%%%%%%%%%%%%%%%%% MODEL INPUT PARAMETER %%%%%%%%%%%%%%%%%%%%%%%%% \newpage | \label{conclusion} %\input{discussion_conclusion.tex} We have presented an extensive study of the formation, orbital dynamics, tidal evolution, and habitability of the \starname system. In Section 2, we presented a simple end-to-end analysis of the accretion of the system. Using the planets' orbital configuration, we built two minimum-mass disks. We then attempted to reproduce the system's orbital architecture. We performed simulations of in situ accretion from these disks, which we interpret as having been shaped by a previous episode of orbital migration. The mass--orbital radius distribution of our simulations provided a modestly good match to the real system, although neither set of simulations adequately matches both the four inner planets and the outer one. The planets also tended to have inclinations that were too large to be consistent with five planets in transit. Perhaps most striking is that our accretion simulations systematically formed too many planets. At least one, and often two, planets tend to form between the orbits of \planete and \planetf. From our dynamical simulations (Section \ref{effectextraplanet}), we can infer that if such a planet exists, then it should be less massive than $1~\mearth$, otherwise its gravitational influence on \planetf would most likely prevent \planetf from transiting. Given that the system is probably older than a few gigayears, simulations of tidal evolution show that the four inner planets of the system are in pseudo-synchronous rotation (respectively, P$_{\rm rot}$ = $\sim$4, $\sim$7, $\sim$13, $\sim$22, and $\sim$130~days) with very low obliquities ($<1\deg$). However, in a few simulations, the obliquity of \planetd was excited to more than 10$\deg$ due to a brief but deep crossing of the 5:3 mean motion resonance between \planetc and \planetd. The competition between the excitation due to planet--planet gravitational interactions and tidal damping has the effect of stabilizing this relatively high obliquity on $\sim$10~Myr timescales. We showed that the eccentricities of the planets cannot be as high as the upper value given by \citet{Quintana2014}. The maximum possible initial eccentricities depend of course on the mass of the planets (through their compositions). A system with planets made of 100\% iron can be stable over 1~Myr for small eccentricities ($\lesssim 0.04$). A system with Earth-composition planets can be stable over 1~Myr with higher eccentricities ($\sim 0.07$) but this can lead to an excitation of the inclinations inconsistent with a transit configuration. A system with 100\% ice planets can be stable over 1~Myr with eccentricities as high as 0.1 for the four inner planets and 0.4 for \planetf. Constraining the mass of the planets would be invaluable information to further constrain the dynamics of this system. We also showed that given the uncertainties on the age of the star as well as the uncertainties on the composition and tidal dissipation, the rotation state of \planetf is unconstrained. If the system is somewhat younger--1~Gyr old--or if the tidal dissipation of \planetf is lower than that of Earth's, then \planetf could still be in the process of pseudo-synchronization and its obliquity could be high. However, if the system is about 4~Gyr old or the tidal dissipation of \planetf is Earth-like, then \planetf would be pseudo-synchronized with a long rotation period ($\sim130$~days). The variety of spin states of \planetf should then be investigated by exoplanet-climate modelers. Our calculations show that the tidal flux generated when the obliquity of the planet is still high is not sufficient to influence the atmosphere of \planetf. The dynamics of the system would be affected if, as predicted by the accretion simulations, an additional planet existed between \planete and \planetf. Without this additional planet, \planetf is relatively isolated from the inner system, so its eccentricity and obliquity oscillations have a very low amplitude. However, an extra planet allows angular momentum to transfer from the inner parts of the system to \planetf, causing higher amplitude oscillations of eccentricity and obliquity. In Section \ref{habi} we presented atmospheric model calculations which indicate that \planetf is indeed squarely situated in the HZ around \starname, with relatively modest amounts of CO$_2$ and N$_2$ required to support conditions conducive to surface liquid water. | 14 | 4 | 1404.4368 |
1404 | 1404.7112_arXiv.txt | We investigate observational constraints on the Brans-Dicke cosmological model using observational data coming from distant supernovae type Ia, the Hubble function $H(z)$ measurements, information coming from the Alcock-Paczy{\'n}ski test, and baryon acoustic oscillations. Our analysis is based on the modified Friedmann function resulting form dynamical investigations of Brans-Dicke cosmology in the vicinity of a de Sitter state. The qualitative theory of dynamical systems enables us to obtain three different behaviors in the vicinity of this state. We find for a linear approach to the de Sitter state $\om=-0.8606^{+0.8281}_{-0.1341}$, for an oscillatory approach to the de Sitter state $\om=-1.1103^{+0.1872}_{-0.1729}$, and for the transient de Sitter state represented by a saddle-type critical point $\om=-2.3837^{+0.4588}_{-4.5459}$. We obtain the mass of the Brans-Dicke scalar field at the present epoch as $m_{\phi}\sim H_{0}$. The Bayesian methods of model comparison are used to discriminate between obtained models. We show that observational data point toward vales of the $\om$ parameter close to the value suggested by the low-energy limit of the bosonic string theory. | Composing the standard cosmological model ($\Lambda$CDM model) we assume that the general relativity describes universe and we postulate validity of the cosmological principle. This model is the best description of the current universe as indicated by implementing the Bayesian methods of model selection to simple theoretical models \cite{Szydlowski:2006ay}. Attempts to explain the present universe in terms of the standard cosmological model is justified by a pragmatic approach of a simple two parameter model. Such a model corresponds to what in physics we know as effective theories, like a standard model in particle physics. In this model as well as in $\Lambda$CDM model there are parameters which value should be obtained from a more fundamental theory or determined by observations. In cosmology the role of such parameters play the density parameters. Unfortunately, the nature of some parameters describing the dark side of the Universe (dark energy and dark matter) is unknown. From our point of view it means that in the construction of the standard cosmological model the cosmological constant term plays only the role of a useful fiction; i.e., the $\Lambda$CDM model describes cosmological observations well but unveils nothing about the nature of the cosmological constant. Adopting the methodology of an effective theory may shed some light on the nature of parameters revealing hints toward a more fundamental theory which we are looking for. Because of well-known problems with the cosmological term in the standard cosmological model related with its substantial interpretation we are looking for a solution of the conundrum of acceleration of the current Universe in the framework of Brans-Dicke theory of gravity \cite{Brans:1961sx} (see also \cite{Clifton:2011jh}). In this theory a gravitational interaction is described in terms of both a scalar field and the metric. The scalar field plays an important role in description of the early universe (inflation) as well as the late time cosmic evolution (quintessence). Moreover following recent Planck observations it is found that the time-varying equation of state with the constant additive contribution is favored when the astrophysical data are taken into account \cite{Ade:2013zuv}. In this framework the difficulty is to obtain a time-varying form of equation of state. In all these applications the scalar field is treated as a source. Usually it is assumed that they are not free and interact with itself via some potential function. Then in scalar field cosmology the main problem is to determine the unknown form of this potential. This problem is passed by assuming different, chosen {\it a priori} forms of the potential function of the scalar field. In the Brans-Dicke theory, which is a scalar-tensor theory of gravity, a scalar field does not play the role of a substance but is rather a integral part of the gravitational sector. In this description a free parameter $\om$ appears as a consequence of the effective theory approach. The value of this parameter can be constrained using the astronomical observations and astrophysical experiments. In the scale of the solar system the Cassini spacecraft mission experiment gave a very stringent bound on $\om > 4000$ for spherically symmetric solutions in the parametrized post-Newtonian (PPN) formalism \cite{Will:book,Will:2005va,Bertotti:2003rm}. On the other hand the data from the cosmological experiments conducted during the WMAP and Planck missions gave substantially lower values of limits on the parameter $\om$. Liddle et al. in \cite{Liddle:1998ij} studied the transition form radiation domination to matter domination epoch in Brans-Dicke theory and showed how the Hubble length at equality depends on the coupling parameter $\om$ for large values of this parameter. Acquaviva et al. in \cite{Acquaviva:2004ti} using structure formation constraints found lower bound $\om>120$ at $95\%$ confidence level. Recently Avilez and Scordis, using CMB data, have obtained the smaller value of the limit $\om > 692$ at a $99\%$ confidence level \cite{Avilez:2013dxa}. Li et al. \cite{Li:2013nwa} using data coming from the Planck satellite and others cosmological observations determined the $\om$ parameter region $-407.0 < \om < 175.87$ at the $95\%$ confidence level, while for positive values of this parameter they obtained $\om > 181.65$ at the $95\%$ confidence level. On the other hand Fabris et al. in \cite{Fabris:2005yt} using the supernovae Ia data obtained the best fit value of $\om = -1.477$. We must remember that all these limits are model dependent. In some estimations the potential of the scalar field is ignored while in others the Newtonian approximation and spherical symmetry is assumed at the starting point. In this paper we find observational constraints on the Brans-Dicke cosmological model assuming the Robertson-Walker symmetry working at the cosmological scale. Therefore the $H^2 (a)$ relation is a starting point of our further estimations of the model parameters. The parameter $\om$ is hidden behind the density parameters of the Brans-Dicke modification of the Friedmann equation. The next step is to estimate the value of the density parameters from the astronomical data and compare the model with the standard cosmological model $\Lambda$CDM using information criteria. Because we treat the new model as a generalization of the $\Lambda$CDM model it is naturally to interpret a prime contribution to the $H^2 (a)$ relation as a corresponding term in the $\Lambda$CDM model. The action for the Brans-Dicke theory \cite{Brans:1961sx} in the so-called Jordan frame is in the following form \cite{Faraoni:book,Capozziello:book}, \begin{equation} \label{eq:action} S = \int\ud^{4}x\sqrt{-g}\left(\phi R -\frac{\om}{\phi}\nabla^{\alpha}\phi\nabla_{\alpha}\phi - 2\,V(\phi)\right) + 16\pi S_m\,, \end{equation} where the barotropic matter is described by \begin{equation} S_{m} = \int\ud^{4}x\sqrt{-g}\mathcal{L}_{m}\,, \end{equation} and $\om$ is a dimensionless parameter of the theory. For the spatially flat Friedmann-Robertson-Walker metric field equations lead to the energy conservation condition \begin{equation} \label{eq:encon} 3H^{2} = \frac{\om}{2}\frac{\dot{\phi}^{2}}{\phi^{2}}+\frac{V(\phi)}{\phi} - 3H\frac{\dot{\phi}}{\phi} + \frac{8\pi}{\phi}\rho_{m}\,, \end{equation} where a dot denotes differentiation with respect to the cosmic time, and the acceleration equation \begin{equation} \label{eq:accel} \begin{split} \dot{H} = & -\frac{\om}{2}\frac{\dot{\phi}^{2}}{\phi^{2}}-\frac{1}{3+2\om}\frac{2V(\phi)-\phi V'(\phi)}{\phi} + \\ & + 2H\frac{\dot{\phi}}{\phi}-\frac{8\pi}{\phi}\rho_{m}\frac{2+\om(1+w_{m})}{3+2\om}\,, \end{split} \end{equation} while the equation of motion for the scalar field is in the following form: \begin{equation} \ddot{\phi}+3H\dot{\phi}=2\frac{2V(\phi)-\phi V'(\phi)}{3+2\om}+8\pi\rho_{m}\frac{1-3w_{m}}{3+2\om}\,. \end{equation} | As we mentioned in the Introduction, the Cassini spacecraft mission in the parametrized post-Newtonian (PPN) formalism gave the most stringent experimental limit $\om>40000$ on the value of the Brans-Dicke parameter \cite{Bertotti:2003rm}. This was obtained in the solar system test for spherically symmetric solutions. The cosmological constraints on the Brans-Dicke parameter $\om$ concern different spatial and temporal scales and the cosmography now plays the role of the PPN formalism. In order to obtain the Hubble functions we did not assume any specific form of the potential function for the Brans-Dicke scalar field. The chameleon mechanism \cite{Khoury:2003aq, Khoury:2003rn, Gubser:2004uf, Brax:2004qh, Mota:2006ed, Mota:2006fz} leads to modifications in the effective potential function, i.e., the effective mass of the scalar field, which depends on the local matter density. In regions of low-mass density like on the cosmological scales, the scalar field is light, while in regions of high density in the solar system, it acquires a large mass, making its effects unobservable. However, the chameleon mechanism is not a generic feature for arbitrary scalar field potential functions. The question whether this mechanism arises for all possible potential functions under considerations remains open. From a theoretical point of view there are two special values of the Brans-Dicke parameter, namely $\om=0$ and $\om=-1$. In the metric formulation of $f(R)$ theory of gravity the action integral, \begin{equation} S=\int\ud^{4}x\sqrt{-g}\,f(R)\,, \end{equation} can be rewritten in the following form \cite{Capozziello:2011et} \begin{equation} S=\int\ud^{4}x\sqrt{-g}\big(\phi R-2V(\phi)\big)\,, \end{equation} which is equivalent to the Brans-Dicke theory with $\om=0$. From the other hand, the Lagrangian density of the low-energy limit of the bosonic string theory \cite{Green:book,Fradkin:1985ys,Tseytlin:1988rr} can be presented in the following form, \begin{equation} \mathcal{L} = e^{-2\Phi}\big(R + 4\nabla^{\alpha}\Phi\,\nabla_{\alpha}\Phi - \Lambda\big)\,, \end{equation} where $\Phi$ is the dilaton field. Making the substitution $\phi=e^{-2\Phi}$, one obtains the Brans-Dicke theory with $\om=-1$ and $V(\phi)=\Lambda\phi$. Neglecting the matter, the two theories are identical, but they differ in their couplings of the scalar field to the other matter \cite{Kolitch:1994qa}. In this paper we obtained cosmological constraints on the models resulting from dynamical analysis of the Brans-Dicke theory. We have shown that for an arbitrary potential function of the Brans-Dicke scalar field, there exists the de Sitter state and that the dynamical behavior in its vicinity crucially depends on the value of the first and second derivative of the scalar field potential function at the de Sitter state as well as on the value of the Brans-Dicke parameter. We found the following mean values of the parameter of the theory: for a linear approach to the de Sitter state $\om=-0.8606^{+0.8281}_{-0.1341}$, for an oscillatory approach to the de Sitter state we obtain $\om=-1.1103^{+0.1872}_{-0.1729}$ while for the transient de Sitter state represented by a saddle-type critical point we find $\om=-2.3837^{+0.4588}_{-4.5459}$. It is interesting that for the models under investigation, for an arbitrary scalar field potential function and excluding the model with $\om<-3/2$ as one leading to ghost behavior, we obtained a value of the $\om$ parameter close to the value needed to obtain correspondence with the low-energy limit of the bosonic string theory. | 14 | 4 | 1404.7112 |
1404 | 1404.0168_arXiv.txt | In this paper, we propose a new method to calculate the mode functions in the noncommutative power-law inflation model. In this model, all the modes created when the stringy space-time uncertainty relation is satisfied are generated inside the Hubble horizon during inflation. It turns out that a linear term describing the noncommutative space-time effect contributes to the power spectra of the scalar and tensor perturbations. Confronting this model with latest results from \textit{Planck} and BICEP2, we constrain the parameters in this model and we find it is well consistent with observations. | By generating an equation of state with a significant negative pressure before the radiation epoch, inflation \cite{Guth:1980zm,Linde:1981mu,Albrecht:1982wi} solves a number of cosmological conundrums, such as the horizon, monopole, entropy problems. After almost thirty-five years of extensive research, inflation is now considered to be a crucial part of the cosmological history of the universe, having affected indelibly its observational features. In the simplest inflation model, the early inflating universe is driven by a scalar field called inflaton, which is classically rolling down the hill of its potential. Inflation predicts a nearly scale-invariant primordial scalar perturbation, which is regarded as the seed of the large scale structures in present. Furthermore, there could be also a primordial tensor perturbation during the inflation time in principle , which is a signal of testing the primordial gravitational waves. All of these are essentially from the quantum fluctuations of the scalar field and the curvature of the universe \cite{Feng:2009kb, Feng:2010ya,Cai:2007et}. These small fluctuations are amplified by the nearly exponential expansion, yielding the scalar and tensor primordial power spectra, which can be observed by measuring the Cosmic Microwave Background (CMB), such as the satellite-based Wilkinson Microwave Anisotropy Probe (WMAP) \cite{Hinshaw:2012aka} and \textit{Planck} \cite{Ade:2013uln} experiments. Although the observed CMB temperature fluctuations, which are generated by scalar perturbations, already helped us to constrain many inflation models, there are still many compelling models that predict almost the same parameter values, which are consistent with observations. A large number of current CMB experiment efforts now target B-\textit{mode} polarization, which could be only generated by tensor perturbations. Recently, a ground-based ``Background Imaging of Cosmic Extragalactic Polarization'' experiment has reported their results (BICEP2). They have shown that the observed B-\textit{mode} power spectrum at certain angular scales is well fitted by a lensed-$\Lambda$CDM + tensor theoretical model with tensor-to-scalar ratio $r=0.20^{+0.07}_{-0.05}$, and $r=0$ is disfavoured at $7.0\sigma$ \cite{Ade:2014xna}. In fact, although there are many inflation models, we still do not known what is the inflaton field. As a candidate for the theory of everything, string theory should tell us how a successful theory of cosmology can be derived from it. General relativity might break down due to the very high energies during inflation, and corrections from string theory might be needed. In the non-perturbative string/M theory, any physical process at the very short distance take an uncertainty relation, called the stringy space-time uncertainty relation (SSUR): \begin{equation}\label{equ:ssur} \Delta t_p \Delta x_p \geq l_s^2 \,, \end{equation} where $l_s$ is the string length scale, and $\Delta t_p = \Delta t$, $\Delta x_p$ are the uncertainties in the physical time and space coordinates. It is suggested that the SSUR is a universal property for strings as well as D-branes \cite{yone, Li:1996rp,Yoneya:2000bt}. Unfortunately, we now have no ideas to derive cosmology directly from string/M theory. Brandenberger and Ho \cite{Brandenberger:2002nq} have proposed a variation of space-time noncommutative field theory to realize the stringy space-time uncertainty relation without breaking any of the global symmetries of the homogeneous isotropic universe. If inflation is affected by physics at a scale close to string scale, one expects that space-time uncertainty must leave vestiges in the CMB power spectrum\cite{Huang:2003zp, Tsujikawa:2003gh,Huang:2003hw,Huang:2003fw,Liu:2004qe,Liu:2004xg,Cai:2007bw, Xue:2007bb}. In this paper, we shall study the power-law inflation in the noncommutative space-time with a different choice of the $\beta_k^\pm$ functions defined below. In this model, it is much more clear to see the effect of noncommutative space-time and much easier to deal with the perturbation functions. A linear contribution to the power spectra of the scalar and tensor perturbations is given in this model. We also confront this model with latest results from the \textit{Planck} and BICEP2 experiments, and we find this model is well consistent with observations. This paper is organized as follows. In next section, we will briefly review cosmological perturbation theory in the noncommutative space-time; in Sec.\ref{sec:power} we calculate the power spectra of the power inflation model in the noncommutative space-time, and compare with observations. In the last section, we will draw our conclusions and give some discussions. And also, in the Appendix.\ref{app:beta}, we presents the detail calculations and discussions on the SSUR algebra and $\beta_k^\pm$ functions. | In this paper, we suggest to take the first form of the $\beta_k^{\pm}$ functions, see Eq.(\ref{equ:beta}). By using this form, it is much more clear to see the effect of noncommutative space-time and much easier to deal with the perturbation functions. A linear contribution to the power spectra of the scalar and tensor perturbations is found in this model. In fact, the second form in Eq.(\ref{equ:beta}) could be also taken by simply redefining $k$ to $-k$, and the results will not be changed. The approximation used in the power-law inflation is $\lambda\ll1$, where $\lambda$ is a free parameter describing the noncommutative effect,see Eq.(\ref{equ:approx}). In other words, all the modes created when the stringy space-time uncertainty relation is satisfied are generated inside the Hubble horizon during inflation. It is not necessarily to consider the case that all the modes are generated outside the Hubble horizon, because in this case all the modes have no causality to each other, and then the flat problem in Big Bang theory can not be solved. After confronting the noncommutative power-law model with the latest results from \textit{Planck} and BICEP2, we constrained the parameter $n$ and $\gamma$, see Fig.\ref{fig}. We conclude that the model is well consistent with the observations. Using the amplitude value of the power spectrum from \textit{Planck}, $\mathcal{R}_s (k=0.002 Mpc^{-1}) = 2.215\times 10^{-19}$ \cite{Ade:2013zuv}, we can also estimate the value of Hubble parameter during inflation of \begin{equation}\label{equ:infhubble} \frac{H_*}{M_{pl}} = \pi \sqrt{r \mathcal{R}_s/2 } \approx 4.67 \times 10^{-5} \,, \end{equation} where $r=0.20$ was used. Then, by using the fitting value of $\lambda \approx 0.01 $, we estimate the string scale as $l_s \approx 2.14 \times 10^2 l_p \approx 1.7\times 10^{-30}$cm, which is a little smaller than that in Refs.\cite{Huang:2003hw, Liu:2004qe}. \appendix | 14 | 4 | 1404.0168 |
1404 | 1404.4866_arXiv.txt | Recent measurements from the BICEP2 cosmic microwave background polarization experiment indicate the presence of primordial gravitational waves with surprisingly large amplitude. If these results are confirmed, they point to a discrepancy with temperature anisotropy power spectrum measurements and suggest that extensions to the standard cosmological model may be required to resolve the discrepancy. One intriguing extension is an anticorrelation between tensors and scalars to naturally suppress the temperature power. Here I examine this possibility and show that such a suppression is not possible in the presence of a general form of anticorrelation. | \label{introsec} The standard $\Lam$ cold dark matter (\LCDM) model has been remarkably successful at describing the large scale geometry, content, and thermal history of the Universe. Nevertheless, whispers of discrepancies from the standard model at the largest observable scales have been heard. In particular, there is a deficit of temperature anisotropy power at the largest angular scales in the cosmic microwave background (CMB) (with respect to the best-fit six parameter model)~\cite{planckresults}. In addition, a roughly dipolar power asymmetry is present on multipole scales $\ell \lesssim 100$ (see, e.g., \cite{planck13isostat}). These features are not of very high significance and are sensitive to {\em a posteriori} choices, so they may simply turn out to be the result of large Gaussian random field fluctuations. Nevertheless, they have attracted considerable attention as they may be hinting at extensions to \LCDM. The whispers of discrepancies have turned into shouts recently with the announcement of results from the BICEP2 CMB polarization experiment~\cite{bicep2}. The BICEP team has performed a measurement of the $B$-mode polarization power spectrum and concluded that their results indicate the presence of a remarkably large-amplitude primordial gravitational wave power spectrum, with tensor-to-scalar ratio $r = 0.16^{+0.06}_{-0.05}$. This result must pass a number of tests before this conclusion can be widely accepted. First, it must be determined whether polarized galactic foreground emission or other systematic effects might account for the signal (see, e.g., \cite{lms14}). Next, the possibility that the signal, if extragalactic, is due to something other than primordial gravitational waves, such as defects, magnetic fields, or birefringence, must be considered~\cite{mp14,lizarragaetal14,bdm14,lln14}. Nevertheless, the BICEP announcement potentially represents one of the most important discoveries in the history of cosmology and should be taken very seriously. One particularly surprising aspect of the BICEP result is the apparent discrepancy with CMB temperature power spectrum measurements. The \Planck\ satellite recently placed a 95\% upper limit of $r < 0.11$~\cite{planckparams}. Such temperature power limits are based on the characteristic shape of the tensor temperature power spectrum, namely, that of a large-scale plateau which tapers off on scales smaller than $\ell \sim 100$. The absence of any visible large-scale power excess limits the possible tensor contribution to within cosmic variance. The fact that we actually observe a large-scale power {\em deficit} only exacerbates this discrepancy, raising it to approximately the $3\sigma$ level~\cite{smithetal14}. Many approaches to resolving this discrepancy, based on the assumption that the BICEP measurement is correct, are possible. Generically, they require the introduction of extra cosmological parameters which have the effect of suppressing temperature power on large scales, to compensate for a large tensor contribution. Perhaps the simplest possibility is the introduction of a negative running of the primordial scalar power spectrum tilt, as pointed out by the BICEP team themselves~\cite{bicep2}. However, the required running is much larger than that expected in the simplest inflationary models (see, e.g., \cite{planckinfl}). Other possibilities include the addition of an anticorrelated isocurvature component~\cite{ksty14} or of additional neutrino species (see, e.g., \cite{archidiaconoetal14}). Of course, an {\em ad hoc} procedure of suppressing the primordial scalar power on the largest scales is certainly a possibility. While inflationary models with such features have been discussed (see, e.g, \cite{hsss14,hsss14b,adss14}), they require an amplitude and cutoff scale tuned remarkably to coincide with and compensate for the tensor contribution. One intriguing possibility is that of an anticorrelation between tensors and scalars, which might offer the possibility of naturally suppressing temperature power without the need to introduce a scale or an amplitude by hand~\cite{cps14}. Of course, such correlations do not occur in the simplest models of inflation, and so would necessitate the introduction of complications to the basic models (see, e.g., \cite{ghp10}). Nevertheless, it is worth investigating the viability of this approach. In this brief report I attempt to address the question of how well temperature power can be reduced with a general form of tensor-scalar correlation. I calculate the total temperature anisotropy power due to tensors and scalars on large angular scales in the presence of such a correlation. I find that a reduction to the total measured temperature power is not possible, in agreement with a special case analyzed very recently in~\cite{cefw14}. | Note that only the {\em total} temperature power remains unsuppressed according to the result of \eq(\ref{summ0}). Individual modes $\alm$ can be expected to be affected by the tensor-scalar correlations. Thus we generically expect the appearance of statistical anisotropy, exhibited as off-diagonal correlations $\bra\alm^{T*}a_{\ell'm'}^S\ket$. In particular, we expect {\em quadrupolar} anisotropy to be induced, i.e.\ couplings between $\ell$ and $\ell \pm 2$, due to the spin-2 character of the tensor modes. However, there are very tight constraints on the presence of quadrupolar asymmetry in the temperature anisotropies. In particular, the \Planck\ measurements are consistent with zero quadrupolar asymmetry even on scales $\ell < 100$~\cite{planck13isostat}, where the effects of tensor correlations would be important. Indeed, as mentioned in the Introduction, the most notable asymmetry is of dipolar character, which should not arise from a tensor correlation. As a logical, if increasingly baroque, possibility, it is worth mentioning that the tensor-scalar correlation tensor $D^{ij}$, which in this work has been assumed to be a constant, could be allowed to vary spatially. With, e.g., a linear gradient in $D^{ij}$, we might expect that dipolar-type anisotropies could be achieved. (Note that it appears to be difficult to reconcile \Planck\ with BICEP by postulating a gradient in $r$ across our observable volume~\cite{cdjky14}.) In order to exhibit analytical expressions, I used the approximations of the SW effect for scalars and the line of sight contribution for tensors. Although these are good approximations on large scales, the structure of \eq(\ref{TScorngen}) should be general in that an improved treatment of the generation of anisotropies will change the detailed form of the transfer functions ($j_\ell$ for scalars and $f_\ell$ for tensors in my approximation), but will leave the integral $\int d\Om_kD^{ij}e_{ij}^\lam(\kh)\sYlm{\lam}{\ell}{m}(\kh)\Ylm^*(\kh)$ unchanged. Thus the final conclusion that a suppression is not possible in the total temperature power should persist. Finally, follow-up observations by the BICEP team, and forthcoming polarization measurements from the \Planck\ satellite, will be crucial in determining whether extensions to \LCDM\ are indeed needed. New measurements indicating a lower tensor-to-scalar ratio than the BICEP2 value may reconcile temperature and polarization measurements while maintaining the exciting consequences of new physics. In such a scenario the motivation for a suppression of large-scale temperature power may be removed. Only future observations will decide whether the ``shouts of discrepancies'' will be silenced or will lead to a new view of the Universe. | 14 | 4 | 1404.4866 |
1404 | 1404.2278_arXiv.txt | {Last year we argued that if slow-roll inflation followed the decay of a false vacuum in a large landscape, the steepening of the scalar potential between the inflationary plateau and the barrier generically leads to a potentially observable suppression of the scalar power spectrum at large distances. Here we revisit this analysis in light of the recent BICEP2 results. Assuming that both the BICEP2 B-mode signal and the \textsl{Planck} analysis of temperature fluctuations hold up, we find that the data now discriminate more sharply between our scenario and $\Lambda$CDM. Nonzero tensor modes exclude standard $\Lambda$CDM with notable but not yet conclusive confidence: at $\sim 3.8\,\sigma$ if $r=0.2$, or at $\sim 3.5\,\sigma$ if $r=0.15$. Of the two steepening models of our previous work, one is now ruled out by existing bounds on spatial curvature. The other entirely reconciles the tension between BICEP2 and Planck. Upcoming $EE$ polarization measurements have the potential to rule out unmodified $\Lambda$CDM decisively. Next generation Large Scale Structure surveys can further increase the significance. More precise measurements of $BB$ at low $\ell$ will help distinguish our scenario from other explanations. If steepening is confirmed, the prospects for detecting open curvature increase but need not be large.} | The smallness of the cosmological constant has led to the consideration of cosmological models with a large number of metastable vacua, in which our universe would arise from the decay of a false vacuum~\cite{Linde:1984ir,Banks:1984cw,Weinberg:1987dv,Bousso:2000xa,Kachru:2003aw,Susskind:2003kw}. Freivogel, Kleban, Martinez, and Susskind~\cite{Freivogel:2005vv} pointed out that in this setting, one expects the inflaton potential to steepen as it interpolates between the slow-roll plateau, where the structure and flatness of our current universe was generated, and the high potential barrier separating it from our parent vacuum, and suggested that this might lead to some observable effect in the power spectrum. Last year \cite{Bousso:2013uia}, we showed that in slow-roll inflation, steepening produces a very specific signal, a suppression (never an enhancement) of the scalar power at large scales. We noted that this effect can resolve the $2-2.5 \,\sigma$ tension at low $\ell$ in the measurements of $C_{\ell}^{TT}$ by the \textsl{Planck} satellite. Since the \textsl{Planck} anomaly was too weak to provide substantial evidence for our signal, we stressed that our analysis should be regarded as a prediction for future observations. We pointed out that in this scenario a similar power suppression should \textit{not} affect the tensor spectrum in the event that it is observed, and we noted that E-mode polarization data, large scale structure, and a nonzero value of the tensor-to-scalar ratio $r$ all have the potential to enhance the significance of a lack of scalar power at low $\ell$. Recently the BICEP2 experiment has reported a detection of primordial tensor modes, with $r=0.2^{+.07}_{-.05}$ \cite{Ade:2014xna,Ade:2014gua}. The importance of this discovery for early universe cosmology, if confirmed by other experiments in the upcoming year or two, is difficult to overstate. In addition to providing almost incontrovertible evidence for an early inflationary phase of our universe, for the first time we will have direct experimental evidence of physics at energies of order $1\%$ of the Planck scale (or not too far from that if different mechanisms to produce gravity waves are involved~\cite{Senatore:2011sp}). The purpose of this note is to reconsider the main observational aspects of \cite{Bousso:2013uia} in light of the possibility of such large values of $r$. (For theoretical motivation and a discussion of priors we refer the reader to \cite{Bousso:2013uia}.) We will see that a value of $r$ as large as that reported by BICEP2 considerably enhances the significance of the low $\ell$ anomaly, to $3.8-4.0\,\sigma$ if $r=0.2$ and $3.5-3.7\,\sigma$ if $r=0.15$. There has been much recent discussion over the apparent tension between large values of $r$ and the \textsl{Planck} bound of $r<0.11$ \cite{Ade:2013uln}. An appealing theoretical interpretation is that the tension lies between the two experiments taken together and $\Lambda$CDM$+r$, rather than between the two experiments. Obviously this could change if the experimental values for the parameters do not stay where they are. If both the BICEP2 and the \textsl{Planck} results hold up to further scrutiny, then it now seems quite likely that the scalar primordial power spectrum deviates from $\Lambda$CDM in a specific way---suppression at large scales--- that arises rather naturally in a cosmology with many vacua.\footnote{In the BICEP2 release it was suggested that modifying $\Lambda$CDM by including running of the tilt can lessen the tension with \textsl{Planck}. This is somewhat true, but the large running required is inconsistent with simple slow-roll models. More concretely, what is needed to make running work is for the potential to have a surprisingly large third derivative in the vicinity of the value of $\phi$ corresponding to $\ell=700$. The effect needed is so large that it would make inflation not last long enough to get a sufficient number of e-foldings before reheating, unless we also introduce a drastic re-flattening of the potential at $\ell\gg 700$~\cite{Easther:2006tv}. We see no theoretical motivation whatsoever for such a feature, unlike the potential steepening at low $\ell$ that we are considering in this paper. Regardless of this theoretical bias, running does not fit the data as easily as a steepening feature. This is because steepening improves the fit precisely in the low $\ell$ region where there is tension with $\Lambda$CDM, whereas running creates new tension at high $\ell$.} Among the many papers that followed the announcement of the BICEP2 results while this work was in progress, some overlap with ours, see for example~\cite{Miranda:2014wga,Smith:2014kka,Hazra:2014aea,Hazra:2014jka}. In particular, the connection to our earlier work \cite{Bousso:2013uia} was noticed in \cite{Hazra:2014aea,Hazra:2014jka}. Earlier related work includes \cite{Linde:1998iw,Garriga:1998he,Linde:1999wv,Contaldi:2003zv,Yamauchi:2011qq,Liddle:2013czu,Dudas:2012vv,Pedro:2013pba}. | The most robust implication of false vacuum decay is negative curvature \cite{Freivogel:2005vv}. But unfortunately, diluting curvature is one of the fortes of inflation. We argued in \cite{Bousso:2013uia} that even if we discover a low-$\ell$ power suppression, and even if it originates from the flanks of the potential barrier separating us from our parent vacuum, we should not necessarily expect that $\Omega_K$ exceeds the cosmic variance limit of $10^{-5}$. The reason is that the slow-roll approximation is still valid in the region where the potential is just starting to steepen, so observable power suppression can arise even if there are still many e-foldings going back in time to the beginning of inflation. Whether or not we should expect to see curvature thus depends on the rate of steepening of the potential, which is something that we do not have well-founded theoretical predictions for.\footnote{Seeing steepening does increase our \textit{chances} for detecting curvature. They may still be small, but we should certainly look for it!} In the absence of direct observation of negative curvature, it is tempting to declare that seeing a low-$\ell$ scalar power suppression is really just evidence for a steepening potential, not for bubble nucleation. We are sympathetic to this point of view (after all in some positivist sense the statement is true), but we now make a few comments about this. Independently of whether or not there is a power suppression in the CMB, bubble nucleation has some compelling theoretical properties \cite{Bousso:2013uia}. It extends in a controlled way our understanding of cosmology to an earlier epoch and provides beautiful homogeneous initial conditions for inflation. Moreover, it arises naturally in a landscape setting such as that of string theory, which has the ability to explain other observations such as the smallness of the cosmological constant. It is thus quite reasonable to \textit{assume} that we nucleated in a bubble, and then to ask if any experimental data could give us evidence against this hypothesis. For example, a detection of positive curvature would kill the scenario \cite{Kleban:2012ph}, but so far there is no evidence for this. In this scenario it would also be rather unnatural, however, to find any nontrivial feature in the inflationary potential \textit{other than} a steepening at early times. This does not mean that the scenario predicts observable steepening; after all, even if it exists, the feature might be located too far up the inflaton potential, so that its imprints would lie far outside our horizon. But if we see \textit{some feature} at low~$\ell$, then it should be power suppression.\footnote{Although a power suppression is not unambiguously predicted in this formalism (in the sense of having probability very close to one), the feature is natural, in the sense that we expect extremely large numbers of e-foldings to be unlikely, which implies that the probability for onset in the observable range of comoving scales is not very small~\cite{Bousso:2013uia}.} For example, a power enhancement at low~$\ell$ would have no motivation in a landscape scenario, and its observation would place considerable pressure on the scenario. Conversely, the discovery of a power suppression, if confirmed, means that the scenario passed a nontrivial check. The above discussion becomes more concrete when we compare false vacuum decay to other models for the initial conditions for slow-roll inflation. Consider a class of theories where $\Omega_K=0$ and the field is taken to start out homogeneously somewhere on the potential. Any such theory is automatically at a disadvantage since, unlike bubble nucleation, it cannot explain why the field began that way, so to some extent this is comparing apples and oranges. Nonetheless, let us allow this comparison.\footnote{Perhaps the most natural way to obtain these initial conditions is from chaotic slow-roll eternal inflation, although we argued in \cite{Bousso:2013uia} that this is probably less common in the landscape than false vacuum eternal inflation. } Our question then should be whether or not seeing a low-$\ell$ power suppression should cause us to significantly modify our priors for these two types of theories. We can argue that it should: in the theory with $\Omega_K=0$ there is no particular reason to expect a steepening perturbation at all. For example, it is sometimes argued that ``simple'' potentials like $m^2 \phi^2$ should be expected for large field models, but in these power-law models the scalar power very gradually \textit{increases} as we go to larger and larger scales and there is never any type of sharp feature.\footnote{In stringy constructions of large-field models by contrast, there are typically constraints which prevent the realization of arbitrary large numbers of e-foldings \cite{McAllister:2008hb,Silverstein:2013wua}. Whether or not the number is small enough to get observable steepening or curvature is model-dependent, and has not been studied in detail.} Nonetheless, we can also compare our scenario to models where there is a steepening feature somewhere in the potential, but not a large landscape. In this case detecting steepening leads to a new coincidence problem. Without a statistical distribution of inflaton potentials and anthropic selection along the lines we argued in \cite{Bousso:2013uia}, there is no particular reason to expect this feature to appear between our horizon and the galactic scale. The discovery of a steepening feature does not \textit{prove} that our universe arose from the decay of a false vacuum in a large landscape. But it leaves us with one more feature of our universe that has a straightforward explanation in this setting, and which would not otherwise have been expected. \paragraph | 14 | 4 | 1404.2278 |
1404 | 1404.7031_arXiv.txt | We calculate the effective masses of neutrons and protons in dense nuclear matter within the microscopic Brueckner-Hartree-Fock many-body theory and study the impact on the neutrino emissivity processes of neutron stars. We compare results based on different nucleon-nucleon potentials and nuclear three-body forces. Useful parametrizations of the numerical results are given. We find substantial in-medium suppression of the emissivities, strongly dependent on the interactions. | With the commissioning of increasingly sophisticated instruments, more and more details of the very faint signals emitted by neutron stars (NS) can be quantitatively monitored. This will allow in the near future an ever increasing accuracy to constrain the theoretical ideas for the ultra-dense matter that composes these objects. One important tool of analysis is the temperature-vs.-age cooling diagram, in which currently a few observed NS are located. NS cooling is over a vast domain of time ($10^{-10}$--$10^5$ yr) dominated by neutrino emission due to several microscopic processes \cite{rep}. The theoretical analysis of these reactions requires, apart from the elementary matrix elements, the knowledge of the density of states of the relevant reaction partners and thus the nucleon effective masses. The present report is focused on the problem of the theoretical determination of this important input information and reports nucleon effective masses in dense nuclear matter obtained within the Brueckner-Hartree-Fock (BHF) theoretical many-body approach. % We study the dependence on the underlying basic two-nucleon and three-nucleon interactions and provide useful parametrizations of the numerical results. Finally some estimates of the related in-medium modification of the various neutrino emission rates in NS matter will be given. We begin with a short review of the BHF formalism and the relevant neutrino emission processes, before presenting our numerical results. | \label{s:end} We have computed nucleon effective masses in the BHF formalism for dense nuclear matter, employing different combinations of two-nucleon and three-nucleon forces. Useful parametrizations of the numerical results were provided. The relevant in-medium correction factors for several neutrino emission processes in $\beta$-stable non-superfluid neutron star matter have then been evaluated in a consistent manner. We find in general in-medium suppression of the emissivities, which however depends strongly on the employed interactions, and reflect mainly the current lack of knowledge regarding nuclear TBF at high density. This emphasizes the need of performing and comparing consistent calculations with given sets of two-body and three-body interactions. | 14 | 4 | 1404.7031 |
1404 | 1404.3612_arXiv.txt | The power spectra of the scalar and tensor perturbations in the noncommutative k-inflation model are calculated in this paper. In this model, all the modes created when the stringy space-time uncertainty relation is satisfied are generated inside the sound/Hubble horizon during inflation for the scalar/tensor perturbations. It turns out that a linear term describing the noncommutative space-time effect contributes to the power spectra of the scalar and tensor perturbations. Confronting the general noncommutative k-inflation model with latest results from \textit{Planck} and BICEP2, and taking $c_S$ and $\lambda$ as free parameters, we find that it is well consistent with observations. However, for the two specific models, i.e. the tachyon and DBI inflation models, it is found that the DBI model is not favored, while the tachyon model lies inside the $1\sigma$ contour, if the e-folds number is assumed to be around $50\sim60$. | Inflation \cite{Guth:1980zm,Linde:1981mu,Albrecht:1982wi} that solves a number of cosmological conundrums, such as the horizon, monopole, entropy problems is now a crucial part of the history of our universe. Within a simplest inflation model, a field called inflaton plays a important role during the early time of the universe, which not only drives the universe to nearly exponentially expand, but also generates small fluctuations that is very important for large-scale structure formation in later time. Inflation may be never end, which we call the eternal inflation \cite{Feng:2009kb, Feng:2010ya,Cai:2007et}. From observations of the Cosmic Microwave Background (CMB), like the satellite-based Wilkinson Microwave Anisotropy Probe (WMAP) \cite{Hinshaw:2012aka} and \textit{Planck} \cite{Ade:2013uln} experiments, we can obtain lots of informations about the early universe. One of them is the power spectra of the perturbations. The observed CMB temperature fluctuations mainly generated by scalar perturbations already have constrain many inflation models, but there are still many left, which are all consistent with observations. Large number of current CMB experiment efforts now target B-\textit{mode} polarization, which could be only generated by tensor perturbations. Recently, a ground-based ``Background Imaging of Cosmic Extragalactic Polarization'' experiment has reported their results (BICEP2). They have shown that the observed B-\textit{mode} power spectrum at certain angular scales is well fitted by a lensed-$\Lambda$CDM + tensor theoretical model with tensor-to-scalar ratio $r=0.20^{+0.07}_{-0.05}$, and $r=0$ is disfavoured at $7.0\sigma$ \cite{Ade:2014xna}. On the other side, general relativity might break down due to the very high energies during inflation, and as a candidate for the theory of everything, string theory should tell us what is a successful theory of the cosmology at that time, and some corrections from string theory might be needed. In the non-perturbative string/M theory, any physical process at the very short distance take an uncertainty relation, called the stringy space-time uncertainty relation (SSUR): \begin{equation}\label{equ:ssur} \Delta t_p \Delta x_p \geq l_s^2 \,, \end{equation} where $l_s$ is the string length scale, and $\Delta t_p = \Delta t$, $\Delta x_p$ are the uncertainties in the physical time and space coordinates. The SSUR may be a universal relation for strings and D-branes \cite{yone, Li:1996rp,Yoneya:2000bt}. Brandenberger and Ho \cite{Brandenberger:2002nq} have proposed a variation of space-time noncommutative field theory to realize the stringy space-time uncertainty relation without breaking any of the global symmetries of the homogeneous isotropic universe. If inflation is affected by physics at a scale close to string scale, one expects that space-time uncertainty must leave vestiges in the CMB power spectrum\cite{Huang:2003zp, Tsujikawa:2003gh,Huang:2003hw,Huang:2003fw,Liu:2004qe,Liu:2004xg,Cai:2007bw, Xue:2007bb}. Recently, Feng \textit{et al.} \cite{Feng:2014yja} propose a new power-law inflation model, by choosing a different $\beta_k^\pm$ functions (shall be defined below), which are equivalent to that proposed by Brandenberger and Ho \cite{Brandenberger:2002nq} by the means of integration. But it is much more clear to see the effect of noncommutative space-time and much easier to deal with the perturbation functions. For detail calculations and discussions on it, see the Appendix~A in Ref.\cite{Feng:2014yja}. It has been shown that a large class of non-quadratic scalar kinetic terms can derive an inflationary evolution without the help of potential term, usually referred to as k-inflation \cite{ArmendarizPicon:1999rj, Garriga:1999vw}, e.g. the tachyon \cite{Sen:2002nu, Sen:2002in} and the Dirac-Born-Infeld (DBI) inflation \cite{Alishahiha:2004eh}. In this paper, we shall study the k-inflation model in the noncommutative space-time following the method in Ref.\cite{Feng:2014yja}. A linear contribution to the power spectra of the scalar and tensor perturbations is given in this model. We also confront two specific k-inflation models, namely the tachyon and DBI models, with latest results from the \textit{Planck} and BICEP2 experiments, and we find that the DBI model is not favored, while the tachyon model lies inside the $1\sigma$ contour, if the e-folds number is assumed to be around $50\sim60$. This paper is organized as follows. In next section, we will calculate the power spectra of the k-inflation model in the noncommutative space-time; in Sec.\ref{sec:kinfmodels} we will constrain the parameters in tachyon and DBI models by using the latest observations. In the last section, we will draw our conclusions and give some discussions. | In conclusion, we have studied the k-inflation model in the noncommutative space-time following the method in Ref.\cite{Feng:2014yja}. A linear contribution to the power spectra of the scalar and tensor perturbations is given in this model. We also confront two specific k-inflation models, namely the tachyon and DBI models, with latest results from the \textit{Planck} and BICEP2 experiments, and we find that the DBI model is not favored, while the tachyon model lies inside the $1\sigma$ contour, if the e-folds number is assumed to be around $50\sim60$. We also constrained the parameter $c_s$ and $\gamma$ for a generic k-infaltion models, and find it is well-consistent with observations, see Fig.\ref{fig:fig1}. | 14 | 4 | 1404.3612 |
1404 | 1404.3424_arXiv.txt | The cold dark matter (CDM) model has two unsolved issues: simulations overpredict the satellite abundance around the Milky Way (MW) and it disagrees with observations of the central densities of dwarf galaxies which prefer constant density (core) profiles.One alternative explanation known as the scalar field dark matter (SFDM) model, assumes that the dark matter is a scalar field of mass($\sim 10^{-22}$ eV/$c^2$); this model can reduce the overabundance issue due to the lack of halo formation below a mass scale of $\sim 10^8$M$_{\odot}$ and successfully fits the density distribution in dwarfs. One of the attractive features of the model is predicting core profiles in halos, although the determination of the core sizes is set by fitting the observational data. We perform \textit{N}-body simulations to explore the influence of tidal forces over a stellar distribution embedded in a SFDM halo orbiting a MW-like SFDM host halo with a disk. Our simulations intend to test the viability of SFDM as an alternative model by comparing the tidal effects that result in this paradigm with those obtained in CDM for similar mass halos. We found that galaxies in subhalos with core profiles and high central densities survive for 10 Gyr. The same occurs for galaxies in low density subhalos located far from the host disk influence, whereas satellites in low density DM halos and in tight orbits can eventually be stripped of stars. We conclude that SFDM shows consistency with results from CDM for dwarf galaxies, but naturally offer a possibility to solve the missing satellite problem. | The Lambda Cold Dark Matter ($\Lambda$CDM) paradigm has been very successful in explaining the structure formation on large scales. One of its predictions is a universal density profile for the dark matter halos. Navarro, Frenk \& White (1997, NFW) suggested a simple formula to describe these density profiles, which presents a divergent inner profile ($\rho(r)\propto r^{-1}$) \citep{die05}. Another prediction of the $\Lambda$CDM model is the number of subhalos per unit mass around the host galaxy. Both predictions have been challenged on scales of dwarf galaxies. In fact, a significant fraction of the rotation curves of low surface brightness (LSB) galaxies and dwarf irregular galaxies are better fitted using dark halos with a density core ($\rho (r) \propto r^{0}$) \cite[]{wil04,gen04,rea06,str06,deb08,oh08,tra08,wal11,agn12,pen12, sal12,lor13}. However, the case for cores in Milky Way (MW) satellites is still debated. For instance, \cite{str14} mentioned that the data of the Sculptor dwarf spheroidal are consistent when an NFW dark matter halo is assumed. Moreover, for dwarf galaxies in the field or Andromeda the information is about the dark matter mass and not the density profiles, so a direct determination of the slope in the density profile is not possible. Regarding the prediction in the number of subhalos, it turns out that the standard CDM model overpredicts the number of dwarf satellite galaxies in the MW and M31. This disagreement is usually referred to as the ``missing satellite problem'' \cite[]{kly99,moo99,goe07,sim07,bel10,mac12,gar14b}. Although the detection of ultra faint galaxies within the MW halo has reduced the missing satellite problem (e.g. Simon \& Geha 2007), a recent study by \citet{iba14} of the distribution of satellites around the MW and M31 suggests they have specific alignments forming planes that are not found in current CDM simulations. Independently of this potential issue, the central densities of MW dSph galaxies are required to be significantly lower than the densities of the largest subhalos found in collisionless DM simulations to agree with current data\citep{boy11,gar14}. Indeed, CDM simulations of the Aquarius Project \citep{nav10} suggest that the MW size halos should inhabit at least eight subhalos with maximum circular velocities exceeding $30$ km s$^{-1}$, while observations indicate that only three satellite galaxies of the MW possess halos with maximum circular velocities $>30$ km s$^{-1}$. This discrepancy is known as the ''too big to fail'' problem. It has been argued that the physics of baryons must be included in order to make a fair interpretation of observations on scales of MW subhalos. For instance, mass outflows given by supernova explosions could transform a cusp into a core in some field dSph galaxies at the present time. The missing satellite discrepancy may be explained as a consequence of gas reionization that quenched the star formation in halos with maximum circular velocity less than $20$ km s$^{-1}$, leaving hundreds of small mass halos without stars \citep{boy14}. In principle using gravitational lensing techniques could confirm the existence of these halos. However, there is still no consensus on whether mass outflows and reionization can explain the observed properties of the MW satellite galaxies \citep{pen12,oka08}. Additionally, it seems there could be a numerical code dependence when interpreting the results obtained from simulations \cite[]{sca01,sca13}. More recently, it was noted that CDM predicts massive subhalos with central densities higher than those found in satellite galaxies, meaning that there are DM subhalos that are massive but host no satellite galaxies \cite[]{boy11,ras12,tol12}. All these problems might be related and share a common solution. The way they are correlated usually depends on the dark model paradigm \cite[]{don13,roc13}, or gravity model \cite[]{mil10,mac13}, but a general fact is that solving one of these issues provides clues to the solution for the other issues. It it worth mentioning that although supernovae explosions seem to play a crucial role in forming cores in field dwarf galaxies and more massive systems where the gas is recycled to continue the star formation \cite[]{mas08,bro11,pon12,gov10,gov12,gar13}, it is unclear that the same feedback implementation works in satellite galaxies where the gas content is negligible and their stellar populations are mostly dominated by old stars. In this sense, dark matter models where the core formation is through DM properties and not by the specifics of astrophysical processes are still viable alternative solutions. One of these alternative models is the scalar field dark matter (SFDM) model. The idea was first considered by \citet{sin94} and independently introduced by \citet{guz00}. In the SFDM model the main hypothesis is that the dark matter is a self-interacting real scalar field of a small mass ($m\sim 10^{-22}$ eV/c$^2$) that condensates forming Bose-Einstein condensate (BEC) ``drops'' \cite[]{mag12a, lor12}. We interpret these BEC drops as the halos of galaxies \citep{mat01} such that the DM wave properties and the Heisenberg uncertainty principle stop the DM phase-space density from growing indefinitely. These properties automatically avoid the divergent density (cuspy) profiles in DM halos and reduce the number of small satellites due to the mass cut-off in the power spectrum \cite[]{hu00,mar14}. For this typical mass, it follows that the critical temperature of condensation of the scalar field is T$_\mathrm{crit}\sim m^{-5/3}\sim$TeV, thus, BEC drops can be formed very early in the universe. There have also been numerous studies that analyzed the behavior of the scalar field at large scales \cite[]{sua11,sua13,mat01,mat07,mag12b,hu00,har11,cha11,ber92}, concluding that it reproduces the successes of the CDM model at those scales. One straightforward and universal prediction of the wave properties of this model is that DM halos have core profiles since their initial formation \cite[]{rob13,sua13}. If halo distributions are flat from the beginning, then strong feedback blowouts are not required to produce low density distributions in DM-dominated systems. Some other consequences of this particular feature have been explored in different contexts; to fit rotation curves in LSB and dwarf galaxies \cite[]{rob13,har11,cha11,lor12,lor14}, and to make strong lensing analyses \cite[]{rob13b, gon13}. All these successes of the model have motivated us to test the model further in order to know if it can be regarded as a serious DM candidate in the universe; conducting these tests is necessary for models whose DM properties are quite different from those of the standard classical particle description. Thus, here we study the evolution of the stellar component of satellite dwarf galaxies embedded in SFDM halos orbiting within an SFDM MW-size host halo. Our study provides constraints on both the final stellar distribution of dSphs and the survival of faint or ultra faint systems. We pursue this task by studying the conditions under which tidal disruption may occur in the SFDM model. Previous studies have shown, using empirical core-like density profiles for DM halos, that tidal disruption can be more important than in halos with NFW profiles, especially if they pass close to the galactic disk (see Klimentowski et al. (2009) and Pe\~narrubia et al. (2010) for collisionless simulations). However, until now there have not been studies addressing whether the tidal effects are strong enough to completely remove the stars in classical and ultra faint dwarf galaxies hosted by BEC halos. The present work aims to investigate this issue through a series of simulations of a stellar component described by a Plummer profile when it is embedded in a SFDM subhalo subject to the influence of a SFDM host halo with a disk component. We also conducted simulations without the disk to compare its effect on the stars. The article is organized as follows: in Section 2 we explain the SFDM model and present the density profiles to be used in the simulations. Section 3 describes the simulations, Section 4 contains our results and discussions of the satellite galaxy evolution, and Section 5 presents our conclusions. \section[]{The baryonic components} \subsection{The dSph stellar component} \label{sec:dwarf} The dSph galaxies have low luminosities ($L_V\sim10^2-10^7$L$_{\odot}$) and very large dynamical mass-to-light ratios $M/L\gtrsim10$, which translate into a large amount of DM \citep{mun05,str06,gil07, mat08,wal09,mac10,wol10}. Nevertheless, we detect the galaxies because of the stars and, in fact, using them as tracers of the potential gives us information about their potential well. Here we use a Plummer density profile \citep{plu11} for the stellar component of the dSph, where the mass density profile is given by \begin{equation} \rho(r)= \frac{3M_{*}}{4 \pi r_{p}^3} \left( 1+\frac{r}{r_{p}} \right) ^{-5/2} \hbox{ ,} \end{equation} where $M_{*}$ is the mass of the stellar component, and $r_p$ is the Plummer radius. One should note that $r_p$ can be related to the half-mass-radius $r_h$ through $r_{h}=1.3r_{p}$. In our simulations, we have set a half-mass-radius of $200$ pc, and a stellar mass of $M_{*}\simeq7.3\times10^5$~M$_{\odot}$, motivated by the typical values for Draco, which is one of the classical dSph galaxies and also one of the least luminous satellites (e.g., see \citeauthor{mar08} \citeyear{mar08} and \citeauthor{oden01} \citeyear{oden01}, for Draco). \subsection{The MW Disk Component} \label{sec:mw_disc} In some of our simulations we include the potential of the MW's baryonic disk, which we model with a Miyamoto-Nagai potential \citep{miy75} \begin{equation} \Phi_d(R,z) = -\frac{G M_d}{\sqrt{ R^2 + (a + \sqrt{z^2+b^2})^2}} \hbox{.} \end{equation} In the latter equation, $M_d$ stands for the mass of the disk, and $a$ and $b$ stand for the horizontal and vertical scalelengths, respectively. We have set the mass of the disk to $M_d=7.7\times10^{10}$~M$_{\odot}$, and the scalelengths to $a=6$~kpc, and $b=0.3$~kpc. \subsection[]{The dark matter component} \label{sec:dm_dwarf} \subsubsection{The dSph DM component} \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_1.eps} \caption{The circular velocity(left), mass(center), and density profile(right) associated to the two SFDM halos of the dwarf galaxy (dashed black line (blue in the online version) is model A: $0.16$~M$_{\odot}$~pc$^{-3}$, and dashed gray line (cyan in the online version) is model B: $0.031$~M$_{\odot}$~pc$^{-3}$). The corresponding core DM models are shown with solid lines with their respective colors for comparison.} \label{fig:halos} \end{figure*} If DM is composed of scalar particles with masses $m \ll$ 1 eV/c$^2$, the galactic halos have very large occupation numbers and the field behaves as a classical field that obeys the Klein-Gordon equation. For SFDM halos, the Newtonian limit is enough to describe them. From the fits to the rotation curves of DM-dominated systems, it has been found that the SFDM halos of DM-dominated galaxies are well described with the ground state \citep{guz06,boh07,rob12,mart14a}. However, the larger the galaxy the more important are the effects of the non-condensed states on the mass profile. The latter means that galaxies that have RC that remains flat even at large radii are better described by adding the excited state contributions \cite[]{ber10,rob13}. This suggests that excited states are relevant to describe MW size systems. Their relevance in dwarfs is out of the scope of this work but see, for instance, \cite{mart15}. In this work we will then consider the base state to describe the dwarf DM halos. Following the hypotheses mentioned above for the SF and using the temperature corrections to one loop for the scalar field, Robles \& Matos (2013) found that after the phase transition that happens in the early universe, the field rolls down to a new minimum of the potential and reaches those values where it will remain. The structures will grow and eventually form the SF halos. Assuming the field is at the minimum, the authors derive an analytical solution for a static spherical configuration that allows the presence of excited states\footnote{We refer the reader to the mentioned work for details on the calculation of the density profile of a SFDM halo given that the mathematical details are already described in that work.}. What they found is that for a SFDM halo in the state $j$ its density profile is given by \begin{equation} \label{eq:sfdm} \rho_j(r)=\rho_{0,j}\frac{\sin^2(k_j r)}{(k_j r)^2} \hbox{ .} \end{equation} In the latter equation, $\rho_{0,j}$ is the central mass density, $k_j \equiv j \pi / R_h$, $j$ is a positive integer that identifies the minimum excited state needed to fit the data of a galaxy, $R_h$ is a scalelength that is determined from observations and its a free parameter; fitting data for a given galaxy provides values for both parameters, the scale $R_h$ and the central density, and the same occurs using eq.(\ref{eq:core}) but for its own parameters. There is not observational evidence that determines how far the halo should extend, but we do expect that halos spread at least enough to cover up to the outermost measured data. It follows that if galaxies have stellar distributions mostly concentrated in the center with possibly some gas surrounding them, then $R_h$ would be larger than the radius where most of the stars are confined. Based on the trend from the fits of works that use scalar field dark matter halos\cite[]{rob12,rob13,lor12,har11} where the scale radius complies with the above condition, we may take the $R_h$ to be a truncation radius such that $\rho(r)=0$ for $r$>$R_h$ and for all $r\leq$ $R_h$ the density is given by eq.(\ref{eq:sfdm}). We recall that scalar field configurations in excited states are characterized by nodes, thus, for a bounded configuration the ground state has no nodes and corresponds to $j=1$, the first excited state has one node and it is associated with $j=2$, the second has two nodes, and so on. We remark that from this interpretation, if we are dealing with a field configuration in the ground state corresponding to zero nodes, then eq.(\ref{eq:sfdm}) has no oscillatory behavior; only those configurations in excited states have oscillations. From equation (\ref{eq:sfdm}) we obtain the mass and rotation curve velocity profiles given by \begin{eqnarray} M(r) &=& \frac{4 \pi \rho_{0,j}}{k_j^2} \frac{r}{2} \biggl(1-\frac{\sin(2 k_j r)}{2 k_j r} \biggr), \\ V^2(r) &=& \frac{4 \pi G \rho_{0,j}}{2 k_j^2} \biggl(1-\frac{\sin(2 k_j r)}{2 k_j r} \biggr) \label{vel}. \end{eqnarray} respectively. \citet{diez14} reported that MW dSphs that are within SFDM halos in the ground state are well described with truncation radii in the range $\sim$ 0.5$-$2 kpc, they mentioned that a common value larger than 5 kpc is disfavored by the dynamics of dSphs provided they are in SFDM halos where only the ground state is taken into account. \cite{mart15} extended this result to account for higher energy states of the scalar field in the associated SFDM halos of the MW dSphs and found that the the stellar distribution lies inside the region where the ground state of the SF is mostly confined ($ \sim$ 0.5 -1.5 kpc), also, the presence of the first excited state does not substantially affect the innermost dark matter configuration but it does allow for the possibility of a larger truncation radius ($\sim$ 5 kpc), nevertheless current data in the dwarfs analyzed are insufficient to conclude the existence of other states in SFDM halos of MW dSphs and hence determining precisely the halo radius. For our generic analysis of a typical dwarf we will then consider a value of $R_h$ =2 kpc consistent with the results suggested by the above independent analyses, additionally, notice that most of the stellar component resides inside 1 kpc in most dSphs, where the dark matter distribution is not substantially modified by the precise halo radius that is considered as shown in \cite{mart15}. For the parameters of the dwarf DM halo we then adopt the values $j=1$ and a typical radius of $R_h=2$ kpc. Notice that for the base state $j=1$ there is no oscillatory behavior in the RC (Figure 1) contrary to what the case with excited states (Figure 2). For the dwarf central density ($\rho_{0,1}$), we select two different values that encompass the range of masses found in dwarfs, $0.16$~M$_{\odot}$~pc$^{-3}$ (model A) and a less massive one with $0.031$~M$_{\odot}$~pc$^{-3}$ (model B). In Figure \ref{fig:halos}, the dashed lines show the circular velocity, mass, and density associated with the SFDM halos of models A (black (blue online)) and B (gray (cyan online)). The corresponding SFDM dwarf core radius (defined as the radius at which the central density drops a factor of two) is $\sim750$~pc for both A and B models and its presence is distinctive prediction of the model. To compare the SFDM profiles with other cored profiles, we also consider the following profile \citep{pen10} \begin{equation} \label{eq:core} \rho(r)=\frac{\rho_0}{(1+(r/R_{s})^2)^{3/2}} \hbox{ .} \end{equation} For both A and B models, we set the scale radius $R_s=1$~kpc (see solid lines of Figure~\ref{fig:halos}). For our mass models, the mass of the dark halo enclosed at $R_{h}=2$ kpc lies in the range $10^{8}$-$10^{9}$ M$_{\odot}$. The resulting mass-to-light ratios represent DM dominated dSphs. For instance, the mass-to-light ratios of dSphs ([M/L]$_{half}$) in the MW range from $\sim 7$~M$_{\odot}$/L$_{\odot}$ (Leo I, Fornax) to $\sim(10^3)$~M$_{\odot}$/L$_{\odot}$ (UMa II, Seg, UMaI) \citep{col14}. In particular, Draco has a very low luminosity but a high estimated total mass within the tidal radius of $M(r_t)=2.2-3.5\times10^7$~M$_{\odot}$ \citep{oden01}, this leads to a high mass-to-light ratio of $(M/L)_{i}\simeq92-146$\citep{irw95,ama95}. \subsubsection{The MW DM component} \label{sec:dm_mw} \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_2.eps} \caption{From left to right are the cicular velocity, mass, and density profiles used for: the MW's SFDM halo model (black lines (purple in the online version)), and the cored DM halo (gray lines (pink in the online version).} \label{fig:haloMW} \end{figure*} In the SFDM model, the fluctuations are expected to grow faster than in the standard model \citep{sua11}, implying that galaxies are fully formed at large redshifts. In fact, some recent high redshift observations suggest the existence of well formed galaxies very early in the universe \citep{col09,fin13,cha14}. From the results of hydrodynamical CDM simulations that model the MW, one sees that its dark and luminous matter do not substantially change since $z \lesssim 2$ ($\sim 10$~Gyr ago) \citep{die07,gov07,kli09,pen10,kas12}. Then, for the initial conditions, we can assume a host with similar parameters that reproduce current MW data. We found that using $\rho_{0,4}=0.0191$~M$_{\odot}$~pc$^{-3}$, $j=4$, and $R_{h}=100$~kpc (see thick solid lines (purple in the online version) in Figure~\ref{fig:haloMW}) in Equation~(\ref{eq:sfdm}) gives a good representation to the MW DM in the SFDM model. Although a detailed analysis of the MW with the SFDM is out of the scope of this work, we obtained the quoted values following the usual procedure of estimating the parameters that model our neighborhood, that is, we search the parameters consistent with the circular velocity in the solar neighborhood and the Oort constants, we find a velocity $\sim$ 200 km/s at 8.5 kpc and constants A=15.5 km s$^{-1}$ kpc$^{-1}$ and B=-14.4km s$^{-1}$ kpc$^{-1}$ similar to previous works\citep{feast97}, we did the estimation when the disk is present and obtained the above values for the SFDM halo and the disk parameters reported following eq. (2). For the MW's cored DM profile (Equation~\ref{eq:core}), we set $R_{s}=15$~kpc. The corresponding circular velocity, mass and density of the cored DM halo, are shown with gray lines (pink in the online version) in Figure~\ref{fig:haloMW}. It has to be noted that, for both (SFDM and cored) MW halos, the core radius is $\sim11.5$~kpc, and that the mass estimations within 100 kpc are comparable. Therefore, the DM profiles are not identical but the total mass enclosed at the halo radius is the same. Given that all our satellites have orbits inside this radius, we choose $R_h$=100 kpc, principally because our main focus is to study the tidal stripping of the stellar component of these satellites, whose apocenters never become larger than 100 kpc during the simulation, as the MW dark matter mass outside this radius is not essential to our study we can truncate the halo at that point, it then follows a Keplerian decay for larger radii. The wiggles found in the halo and shown in Figure \ref{fig:haloMW} are also a particular difference of this SFDM profile with respect to other core models. \section[]{Simulations} \label{sec:code} We simulate the evolution of the stellar component of the dwarf galaxy, which is embedded in a rigid SFDM halo potential using the $N$-body code \scriptsize {SUPERBOX} \normalsize \citep{fellhauer00}. \scriptsize {SUPERBOX} \normalsize is a highly efficient particle-mesh, collisionless-dynamics code with high resolution sub-grids. In our case, \scriptsize {SUPERBOX} \normalsize uses three nested grids centered in the density center of the dwarf galaxy. We used $128^3$ cubic cells for each of the grids. The inner grid is meant to resolve the inner region of the dwarf galaxy. The spatial resolution is determined by the number of grid cells per dimension ($N_c$) and the grid radius ($r_{\rm grid}$). Then the side length of one grid cell is defined as $l=\frac{2 r_{\rm grid}}{N_c-4}$. For $N_{c}=128$, the resolution is $0.5$~pc. \scriptsize {SUPERBOX} \normalsize integrates the equations of motion with a leap-frog algorithm, and a constant time step $dt$. We selected a time step of $dt=1$~Myr in our simulations. \begin{table*} \centering \caption{Parameters used in our simulations.\label{table:1}} \begin{tabular}{ccccccc} \tableline\tableline & & Dwarf & Dwarf & MW & MW & DM\\ Simulation & $\frac{r_{p}}{r_{a}}$& orbit & $\rho_{0}$ & $\rho_{0}$ & disk & model \\ & & plane &$(10^7$ M$_{\odot}$~(kpc)$^{-3})$ & $(10^7$ M$_{\odot}$~(kpc)$^{-3})$& &\\ \tableline A1 & 1/2 &x-y& 16 & 1.91 & --& SFDM \\ B1 & 1/2 &x-y& 3.1& 1.91 & --& SFDM \\ A2 & 1 &x-y& 16 & 1.91 &$\checkmark$ & SFDM\\ B2 & 1 &x-y& 3.1& 1.91 &$\checkmark$ & SFDM\\ A3 & 1/2 &x-y& 16 & 1.91 &$\checkmark$ & SFDM\\ A3$_{core}$ & 1/2 &x-y& 16 & 1.91 &$\checkmark$ & Core\\ B3,~B6\tablenotemark{a} & 1/2 &x-y& 3.1& 1.91 &$\checkmark$ & SFDM\\ B3$_{core}$,~B6$_{core}$ & 1/2 &x-y& 3.1& 1.91 &$\checkmark$& Core \\ A4,~A6\tablenotemark{b} & 1/5 &x-y& 16 & 1.91 &$\checkmark$ & SFDM\\ A4$_{core}$,~A6$_{core}$ & 1/5 &x-y& 16 & 1.91 &$\checkmark$& Core \\ B4 & 1/5 &x-y& 3.1& 1.91 &$\checkmark$ & SFDM\\ B4$_{core}$ & 1/5 &x-y& 3.1& 1.91 &$\checkmark$ & Core\\ A5 & 1/5 &45$^{\circ}$& 16 & 1.91 &$\checkmark$ & SFDM\\ B5 & 1/5 &45$^{\circ}$& 3.1& 1.91 &$\checkmark$ & SFDM\\ \tableline \end{tabular} \tablenotetext{a}{Simulations B6(B6$_{core}$) use the same parameters of B3(B3$_{core}$) but with a satellite stellar mass M$_{\ast} = 1 \times 10^4 \ M_{\odot}$}. \tablenotetext{b}{Simulations A6 (A6$_{core}$) use the same parameters of A4(A4$_{core}$) but with a satellite stellar mass M$_{\ast}=1 \times 10^4 M_{\odot}$}. \tablecomments{Column 1 identifies the simulation, column 2 specifies $r_{p}$/$r_{a}$ for the orbit, column 3 shows the plane of the orbit, next two columns give the central density for the dwarf and the MW DM halos in each simulation, respectively, column 6 determines if a disk is present in the Milky Way halo, and column 7 gives the DM model used in the simulation.} \end{table*} \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_3.eps} \caption{The surface mass density of the dwarf galaxy for models $A1,B1,A2$, and $B2$ for $t=0$, $3$, $6$, and $10$~Gyr, all plots are centered in the dwarf galaxy. In the last column, we show the orbit of the satellite galaxy around a Milky Way SFDM halo (colored version online).} \label{fig:vic1} \end{figure*} For the orbit of the dwarf galaxy, we assume an apocenter distance from the MW, $r_{a}=70$ kpc \citep{bon04} and two different pericenter distances ($r_p=14$ and $35$ kpc). We conducted simulations with and without adding the presence of a Miyamoto-Nagai disk in the MW potential to assess the effects on the dwarfs due to the close encounter with the disk component. Our main interest is the stellar component evolution that is located deep inside the subhalo. \cite{pen10} show that the major effect of tidal disruption of a DM suhalo occurs in the outermost radius, while inner regions ($\lesssim 1$ kpc) are less affected by tides and the density profiles are only shifted to a slightly lower value maintaining the same inner shape during the evolution. The evolution changes if the subhalo's pericenter is smaller than the length of the disk in the event that this component is present, meaning that when the subhalo effectively cross through the disk several times it can lose a considerable amount of its initial mass or even get destroyed if the orbit's pericenter is $\sim$1.8 kpc, however, these are rare events. Given these results and that we consider rigid subhalos, we focus our analyses on orbits with pericenters larger than the disk scalelength avoiding direct collisions with the disk that would require a live subhalo. Since stars serve as tracers of the subhalo potential, any major tidal disruption of the stars would be indicative of a substantial change in the evolution of the subhalo. In such cases, a live halo would be needed. This happens only in one of our simulations and will not be used to draw the overall conclusions of this work. However, it does serve to show that our results are consistent with those in \cite{pen10}. In our first couple of simulations, denoted by $A1$ and $B1$, the dwarf galaxy is embedded in the MW SFDM halo potential without including the baryonic MW disk component (first two rows of Figure~\ref{fig:vic1}). The dwarf galaxy is placed at a galactocentric distance of 70~kpc, and orbits in the $x-y$ plane with a $r_{p}/r_{a}=1/2$. \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_4.eps} \caption{Same as Figure~\ref{fig:vic1}, but for models $A3-B4$(colored version online).} \label{fig:vic2} \end{figure*} In the second pair of simulations, $A2$ and $B2$, we model the dwarf galaxy embedded in the MW SFDM halo potential in a circular orbit ($r_{p}/r_{a}=1$), including the baryonic MW disk component (see last two rows of Figure~\ref{fig:vic1}). In the third pair of simulations, $A3$ and $B3$, we model the dwarf galaxy embedded in the MW SFDM halo potential, with a $r_{p}/r_{a}=1/2$, but now we include a rigid baryonic MW disk component. We rerun these two simulations to compare with the empirical profile (eq. \ref{eq:core}) referred as $A3_{core}$ and $B3_{core}$. We observe from Figures~\ref{fig:vic1} and \ref{fig:vic2} that the dwarf galaxy survives unperturbed for $\sim10$~Gyr in models $A1-B3$. Moreover, from models $A1$, $A3$, $B1$, and $B3$, we observe that there is a negligible effect of the MW's baryonic disk on the dwarfs that are in SFDM subhalos. \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_5.eps} \caption{The surface mass density of the dwarf galaxy for $t=0$, $3$, $6$, and $10$~Gyr, for models $A4$ and $B4$ centered in the MW SFDM halo potential.The white cross shows the center of the MW, and the white line shows the dwarf's orbit around it(colored version online).} \label{fig:vic4} \end{figure*} The fourth pair of simulations, named as A4 and B4, resembles cases 3 but with $r_{p}/r_{a}=1/5$, ($A4$ and $B4$ are in Figure \ref{fig:vic2} and Figure~\ref{fig:vic4}). For model $A4$, the stellar component of the dwarf galaxy remains undisturbed, while the $B4$ model shows a major star mass loss. We run an extra couple of simulations for completeness as discussed in the next section. Table 1 summarizes all our simulations. | Figure~\ref{fig:mass} shows the dwarf galaxy stellar mass profile at $t=0$ and $t$ = 10~Gyr in all our simulations. From the upper left panel in Figure \ref{fig:mass} we see that all $A$ models lose some particles, but the loss is not substantial and the galaxies survive with essentially the same initial mass after $10$~Gyr. These simulations suggest that the density is high enough to strongly bind the stars and prevent the disruption of the satellite. A similar behavior is seen when a cuspy-like profile is used \citep{kli09,lok12}, making tidal disruption an inefficient process in both core and cusp-like subhalos to reduce their stellar mass within 1kpc and therefore making it not the relevant mechanism that decreases the abundance of massive dwarf satellites around MW-type galaxies, even for orbits with close pericenters of 14 kpc. The $B$ models for the SFDM halo show a slightly larger particle loss than the $A$ models (upper right panel in Figure 6) except for model $B4$ which shows a more pronounced particle loss. The small central density of the SFDM dwarf subhalo, plays a crucial role in its survival. The final mass (at $t=10$~Gyr) for $B$ models is smaller than the high $A$ density case in all cases. This shows that even if the orbit is far from the MW disk, whenever the DM subhalos have low densities the stars in the center are susceptible to spread out more than in denser halos as seen by comparing the two upper panels in Figure \ref{fig:mass} within 500 pc. One of the features that is seen from the stellar mass profiles is that the stars are not heavily stripped from the dwarf SFDM halo (excluding model $B4$). This is reassuring as it implies the DM density profile is also not strongly modified in that region and may be approximated by a fixed halo profile for orbits without small pericenters. The result is strengthened with the findings of \cite{pen10} for live subhalos with $r_{p}/r_{a}=1/2$ and a core DM profile; even in the presence of a live disk the DM subhalos remain almost the same in the central region after $10$~Gyr. Hence, the tidal effects on the subhalo are small within $1$~kpc, which is about the relevant core size of our simulated subhalo. Therefore we consider that our approximation of a SFDM rigid halo is sufficient as long as the subhalos do not get well inside the disk of the host halo. \begin{figure*} \centering \includegraphics[width=1.0\textwidth]{FIG_6A.eps} \includegraphics[width=1.0\textwidth]{FIG_6B.eps} \caption{Upper panel: Dwarf stellar mass for models A(left) and B(right) at $t=0$ and $t=10$~Gyr. The upper left panel would represent a classical dwarf and the upper right would be an ultra faint-like galaxy. The different symbols in the panel represent the dark matter model used in that simulation according to Table 1. Bottom panels show small mass satellites with M$_{\ast}=1\times 10^4$ (models A6, A6$_{core}$, B6, and B6$_{core}$) at $t=0$ and $t=10$~Gyr, different symbols correspond to different simulations. In all $A$ models of the SFDM the galaxy survives at the end of the simulation independent of the stellar mass and the orbits we considered, even the presence of a disk in the MW scalar field halo cannot destroy the satellite. In B models where the subhalo is less dense, the satellite losses more mass than in A cases but will still survive inside the subhalo, except when the pericenter becomes comparable to the disc's scale length where we expect the scalar field subhalo to be disrupted too (color version online).} \label{fig:mass} \end{figure*} Simulations $A3_{core}$ and $A4_{core}$ present a similar behavior than their SFDM counterparts (see upper left panel of Fig.~\ref{fig:mass}). In these cases the stars in the outskirts get stripped more easily, moving to larger radii and, at the same time, causing the inner stars to redistribute to a new configuration that follows the background DM halo potential. For cases A, the potential well is deep enough that only a few stars are lost; most of them remain within 1 kpc and keep the same initial profile. The $B3_{core}$ model has lost more mass than its analog $B3$ (see the upper right panel of Fig.~\ref{fig:mass}). This is due to the slight difference in the tail of the subhalo mass profile ($r$> 2kpc) and the fact that the potential is not as deep as in cases $A$, making it easier for the tidal forces to change the central stellar distribution. From Figure 1 we note that subhalos with a core profile have a non zero density for $r$>2 kpc. For smaller $r_p:r_a$ the tidal stripping and the interaction with the disk becomes stronger, especially for the stars in the outermost radius which are more easily stripped. In fact, given that the subhalo in $B3_{core}$ is more extended than in $B3$, more stars are likely to get pulled toward the tail of the halo but remain inside the subhalo. In this process the now outer stars drag some of the inner stars towards outer radii producing a more extended stellar distribution than in $B3$, reducing at the same time the stellar mass as shown in Fig.~\ref{fig:mass}. The same occurs for the $B4_{core}$ model and its counterpart $B4$ (see empty and full squares in Figure~\ref{fig:mass}). However, the considerable disruption in both $B4$ simulations indicates the need to include the disruption of the halo. In our simulations, the satellites still remain due to the assumption of fixed subhalo, but we expect the dwarf halos to fully disrupt and that their stars get dispersed around the MW halo. We also conducted a couple of simulations ($A5$ and $B5$) where we set the dwarf galaxy embedded in the MW SFDM halo potential with a $r_{p}/r_{a}=1/5$ including the baryonic MW disk component and similar to $A4$ and $B4$ models, but now we place the dwarf galaxy in an orbit inclined $45^{\circ}$ from the $x-y$ plane. In this case we observe that the dwarf galaxy gets destroyed within $\sim 1.5$~Gyr. This suggests that orbits with close pericenter distances and inclination effects are factors important to the survival of low density SFDM dwarf satellites. In order to address the dependence of our results on the stellar mass we conducted the following two pairs of simulations. In $A6$ and $A6_{core}$, the parameters are identical to $A4$ and $A4_{core}$ respectively, but the stellar mass of the satellite is smaller M$_{\ast}=1\times 10^4 M_{\odot}$.These parameters correspond to the closest orbit where the tidal effects should be the largest. The other pair, $B6$ and $B6_{core}$, uses $M_{\ast}= 1\times 10^4 M_{\odot}$ and the parameters of $B3$ and $B3_{core}$ respectively. We do not use parameters of $B4$ for the reasons mentioned in the previous paragraphs. These cases are shown in the bottom panels of Fig.~\ref{fig:mass}. In the bottom left of this figure we notice that the inner mass of the satellite reduces due to the proximity to the disk's influence but the potential well is again deep enough to ensure the survival of the satellite. In the bottom right panel of the same figure we see that both cases $B6$ lead effectively to the same result; despite their low masses the satellites can remain with most of their initial mass after 10 Gyr. We point out that this is consistent with the arguments given above for $B3_{core}$. In $B6_{core}$ the satellite has a much smaller initial stellar mass concentrated within $1$kpc, thus fewer stars are stripped during its evolution and are insufficient to drag most of the central stars toward the outer regions as opposed to $B3_{core}$. Nevertheless, we still observe a small effect of this process in this pair of simulations. Indeed, comparing cases $A$ and $B$, we notice that for the cases in which orbits are far from disk the density of the subhalo is a decisive parameter for determining the mass loss but has little influence in its survival. We found that the mass loss is greater if the subhalo has smaller density, but the dwarf galaxies still survive after 10 Gyr. When the satellites have orbits close to the center of the host or when they strongly interact with the disk, the subhalo central density becomes an important factor for the survival and for the number of remaining satellites; this is consistent with previous works. It is known that in CDM simulations, the satellites with cuspy subhalos can be stripped of stars but still survive as DM-only subhalos, which could be detected with gravitational lensing techniques. Here we show that if the satellites are in scalar field subhalos with central densities comparable to classical dSphs, some of their stars are stripped but the galaxies can survive with smaller masses and hence contribute to the number of dwarf satellites around a MW host. It must be noted that, in the SFDM model, the substructure is smaller due to the wave properties causing the cut-off in the power spectrum, as confirmed in \cite{sch14}. Our result can be tested with hydrodynamical SFDM cosmological simulations in the future. For the lower density dwarfs (comparable to ultra faint dwarfs), we obtained that they could survive but only if their orbits do not get well inside the disks of their hosts. On the other hand, low density halos with close pericenter orbits can be fully stripped of stars if evolved for a long time even with a fixed subhalo potential, but as mentioned before we expect them to be destroyed once the fixed halo hypothesis is relaxed. Therefore, we do not get DM halos that are tiny and dark, contrary to the CDM predictions where the cusp prevents total disruption. The formation of ultra faint dwarfs is still not clear but it is thought that they are the result of more massive dwarfs that were disrupted and left them as low density systems. We have seen that in the scenario of SFDM, depending on their distance to their host, these faint systems could also be produced from disrupted dwarfs with initial core profiles in the same way that they are when halos are assumed to have cusp profiles \citep{lok12}. Therefore, our results point to an alternative solution to the satellite overabundance problem and the cusp-core issue by means of the quantum DM properties of the scalar field without relying strongly on the messy astrophysical processes. Here, small mass subhalos with core profiles ($\rho(r) \sim r^{0}$) and with orbits not crossing the host's disk are able to survive for a long time; otherwise the close encounters with the disk could completely destroy them. On the other hand, more massive dwarfs can get closer or farther from the host disk and still survive with core profiles. To determine the final fate of these galaxies and test the results from the present work, we will need simulations that involve the complexities of astrophysics, but we leave that for a future work. From eq. (3) we see that for larger excited states the inner region becomes more compact and the halo core sizes can become smaller. In general a superposition of states may be present in a SFDM halo\citep{rob13}, however, in \cite{ure10} they found that for a stable multi-state SFDM halo the number of bosons in excited states should decrease with increasing $j$, for instance, in a multi-state composed of the ground($j$=1) and first excited states ($j$=2) the number of bosons in $j$=2 should be less than or equal to those in $j$=1 if the halo is to be stable, hence small core sizes are expected in massive galaxies, i.e., those where the SF ground state provides a poor description, depending on the required number of excited states in the halo the core size can become smaller. Given that dwarf galaxies are consistent with halos with only $j$=1, and from to the stability constraint in the SFDM halos any contribution of higher energy states will be subdominant in such halos and the core sizes are dictated by the dominant ground state, being of $\sim$ kpc for a boson mass of $\sim$10$^{-22}$ eV/$c^2$. By fitting galaxies of different sizes it is possible to constrain the number of states in the SFDM halos required to agree with observations. As noted before, dwarfs seem to lie in ground state SFDM halos, while larger galaxies seem to require more than that, however, at this point it is unclear whether there is a mechanism that predicts the final superposition of a given SFDM halo; a statistical analysis fitting galaxies of different morphological types and sizes is required to derive such a relation. Additionally, although we use a single state($j$=4) as a fit to the MW, we have not addressed its stability. Looking at the more general SFDM halo scenario mentioned above, it remains to be shown under what circumstances the interplay of the different particle states within a SFDM halo, in our case a MW-like halo, can lead to a stable halo with a non-negligible fraction of bosons in a single state that extends to large radii; this is something that needs to be addressed at a later time and will be vital to the viability of SFDM. | 14 | 4 | 1404.3424 |
1404 | 1404.6940_arXiv.txt | Giant planets like Jupiter and Saturn feature strong zonal wind patterns on their surfaces. Although several different mechanisms that may drive these jets have been proposed over the last decades, the origin of the zonal winds is still unclear. Here, we explore the possibility that the interplay of planetary rotation with the compression and expansion of the convecting fluid can drive multiple deep zonal jets by a compressional \neu{Rhines-type}\alt{beta} mechanism, as originally proposed by \citet{Ingersoll1982}. In a certain limit, this deep mechanism is shown to be mathematically analogous to the classical \neu{Rhines}\alt{beta} mechanism possibly operating at cloud level. Jets are predicted to occur on a compressional Rhines length $l_R = (2 \Omega \langle H_\rho^{-1} \rangle v_{jet}^{-1} )^{-1/2}$, where $\Omega$ is the angular velocity, $\langle H_\rho^{-1} \rangle$ is the mean inverse density scale height and $v_{jet}$ is the typical jet velocity. Two-dimensional numerical simulations using the anelastic approximation reveal that this mechanism robustly generates jets of the predicted width, and that it typically dominates the dynamics in systems deeper than $O(l_R)$. Potential vorticity staircases are observed to form spontaneously and are typically accompanied by unstably stratified buoyancy staircases. The mechanism only operates at large rotation rates, exceeding those typically reached in three-dimensional simulations of deep convection in spherical shells. Applied to Jupiter and Saturn, the compressional Rhines scaling reasonably fits the available observations. Interestingly, even weak vertical density variations such as those in the Earth core can give rise to a large number of jets, leading to fundamentally different flow structures than predicted by the Boussinesq models typically used in this context. | \label{introduction} Strong zonal winds organize the colorful clouds on Jupiter's surface into banded structures. These cloud patterns have already been observed with telescopes more than 350 years ago \citep{Rogers1995}, and since then, scientists have studied them with ever more sophisticated observation tools. In addition to optical telescopes, the Hubble space telescope and the Pioneer, Voyager, Galileo and Cassini spacecraft missions have revealed fascinating pictures of the complex wind patterns shaping the surfaces of the gas and ice giants in our planetary system. Jupiter and Saturn exhibit strong prograde (i.e. eastward) equatorial jets, which are flanked by weaker, alternating westward and eastward winds on each hemisphere. Uranus and Neptune also exhibit pronounced zonal winds, but in contrast to Jupiter and Saturn, strong retrograde equatorial jets are observed. It is still unknown how deep the winds extent into the interior. The Galileo probe that entered Jupiter's atmosphere down to $150$ km in 1995 gave evidence for an increase of the wind speeds at larger depth \citep{Atkinson1997}, but provided little information about the deep interior. The Juno mission that will reach Jupiter in mid-2016 is expected to better constrain the radial extent of the jets by carrying out high-resolution measurements of Jupiter's gravity field \citep{Kaspi2010}. The theoretical understanding of the zonal winds is still incomplete. The different theories proposed so far are commonly classified into two distinct groups. The first class of models suspects the key to the zonal winds in a shallow layer at cloud level, with energy being pumped into the jets by processes like moist convection, lateral variations in solar heating or other processes occurring close to the surface. In contrast, the second class of models views the zonal winds as an expression of processes occurring deep in the planetary interior, typically driven by convective instabilities. In the following, we will refer to these two classes as shallow- and deep-forcing models. They represent end-member cases of potential forcing scenarios, and it is possible that a combination of both, deep- and shallow-forcing, is needed in order to explain all observational data. It is instructive to reconsider the key processes driving zonal flows in the different approaches. The models following the shallow-forcing paradigm typically consider a fluid confined to a thin layer at the planetary surface, which leads to considerable simplifications of the governing equations. In the simplest cases, two-dimensional, incompressible flow is assumed, while more advanced approaches include shallow-water and multi-layer models. In all these cases, the latitudinal variation of the tangential component of the Coriolis force, the so-called {\em beta effect}, plays a key role, as it forces fluid parcels moved in latitudinal direction to change their vorticity. The effect becomes significant for large flow structures only, whereas the dynamics on small and intermediate scales is usually characterized by an inverse cascade of kinetic energy. This turbulent upward cascade ceases at the so-called Rhines length \citep{Rhines1975} when the beta effect becomes felt by the flow. From this length scale on, the flow dynamics is dominated by Rossby waves, which leads to a strong anisotropy of the large scales and ultimately to the formation of jets. A large number of theoretical, experimental and numerical studies have confirmed the robustness of this now classical picture of zonal wind generation (see \citet{Vasavada2005} for a review). While earlier works typically report retrograde equatorial jets, more recent simulations have also succeeded in producing prograde equatorial jets by including additional physical processes like energy dissipation by radiative relaxation or latent heating resulting from the condensation of water vapor (e.g. \citet{Cho1996,Williams2003,Showman2004,Scott2008,Lian2010}). In contrast to these shallow models, in the deep-forcing scenario, convection in a fluid confined to a deep spherical shell is considered. In case of the gas giants, the inner boundary is often assumed to be set by the transition from molecular to metallic hydrogen, where Lorentz forces become important and are thought to lead to different flow dynamics deeper within the planet. As in the shallow models, Rossby waves are believed to play a central role in driving deeply seated zonal jets as well. However, the processes of local vorticity generation that are responsible for the waves are governed by different physics. Mainly, two mechanisms have been identified as possible sources of Rossby waves in the deep interior -- the so-called {\em topographic beta effect} and a process that we call the {\em compressional beta effect} in this paper. Both are reviewed in more detail below. Apart from different wave mechanisms considered, theories on deep jet generation also differ in their mechanistic view on how these waves channel kinetic energy into zonal winds. Table \ref{mechanism_table} gives a schematic overview over several popular models, along with their underlying assumptions and key predictions. Even though this table only presents a somewhat simplified view and fails to include all aspects of the various theories, we feel that it is nevertheless useful in providing a compact overview of the similarities and differences between the various approaches. Note that these different views of the jet generation mechanics are complementary in many ways, and should not be seen as being necessarily contradicting. In the following, we give a brief review of the different models, which allows us to discuss the ``compressional beta effect'' studied in this paper in a broader context. \begin{table*}[bt] \label{mechanism_table} \begin{center} \includegraphics[width=\linewidth]{mechanism_table} \end{center} \caption{List of popular mechanisms possibly driving zonal winds in giant planets in the deep-forcing scenario. The above mechanisms can be subdivided into two groups that consider different Rossby wave sources, i.e. the topographic- and the compressional beta effect. Within each group, different mechanistic pictures of the jet generation process exist, which are discussed in more detail in the text. Note that these mechanistic views may in many respects be complementary and do not necessarily oppose each other.} \end{table*} The vast majority of studies performed so far are based on the {\em topographic beta effect} as the source of Rossby waves. Since the large scales in rapidly rotating flows are strongly affected by Coriolis forces, coherent columnar structures parallel to the rotation axis can be expected to exist in giant planets. Such fluid columns may extend through the entire planet touching the spherical boundaries. When moved perpendicular to the rotation axis, they undergo vertical stretching or compression due to the spherical geometry. This, through angular momentum conservation, results in a change of their vorticity, a process often called {\em topographic beta effect}. Since the newly generated vorticity is out of phase with the vorticity associated with the outward and inward movement of the fluid columns, azimuthally propagating Rossby waves can be generated. As shown by \citet{Busse1970}, convection is most easily excited outside an imaginary cylinder parallel to the rotation axis enclosing the inner spherical shell, the so-called tangent cylinder. Due to the topographic beta effect, the flow takes the form of propagating convection waves, which are essentially Rossby waves modified by the thermal buoyancy field (so-called {\em thermal Rossby waves}). Because of the curved spherical boundaries, the strength of the topographic beta effect depends on the distance from the rotation axis, which causes a slower wave propagation close to the tangent cylinder than further outside. This in turn leads to a tilt of the convection columns in prograde direction towards the outer boundary, which gives rise to Reynolds stresses transporting prograde momentum outwards and retrograde momentum inwards. A zonal flow is generated, which is reinforced by a mean flow instability that can drive a strong prograde equatorial jet (see e.g. Busse, 2002, for a review). \citet{Busse1983,Busse1994} further argued that multiple zonal jets can be generated because the limited radial extent of the convection columns allows for a breakup of convection into cylindrical sub-layers, such that the differential rotation pattern observed on Jupiter may be explained. In table \ref{mechanism_table}, we refer to this view of differential rotation generation as the {\em Busse picture}. Numerical simulations \citep{Heimpel2005,Heimpel2007,Jones2009} have shown that an alternating jet pattern very similar to the one observed e.g. on Jupiter, can form if the ratio of the inner to outer shell radius is chosen large enough, i.e. if the shell is assumed sufficiently thin. Typically, a strong prograde jet is found at the equator, which is readily explained by the ideas described above. The high latitude jets observed in the simulations are interpreted by the authors as being a consequence of a Rhines-type mechanism, in which energy is transported to the large flow scales by an inverse cascade. As in the shallow models, the upscale transport finally comes to a halt when Rossby waves become important in the dynamics and channel the kinetic energy into zonal jets. The number of the jets observed in the simulations has indeed been shown to follow an approximate Rhines-type scaling. In table \ref{mechanism_table}, we call this view of zonal flow generation the {\em topographic Rhines picture}. The topographic beta effect, i.e. the local vorticity generation due to a change in system height, is not the only process possibly generating Rossby waves in deep planetary atmospheres. If the strong density stratification in giant planets is taken into account, local vorticity can also be produced in another way. As fluid parcels rise and sink in the background pressure field, they expand and contract. Because the Coriolis force exerts a torque on such fluid parcels, vorticity is locally produced or destroyed \citep{Glatzmaier1981,Glatzmaier2009}. This local vorticity source can be interpreted as a beta-type effect \citep{Ingersoll1982}, and since it crucially relies on the compressibility of the planetary gas, we call it the {\em compressional beta effect}. The vorticity generation is again out of phase with the vorticity associated with the upwelling and downwelling fluid, such that Rossby waves can be generated. Different from the topographic beta effect however, the propagation speed of these Rossby waves is proportional to the local inverse density scale height, and is not dictated by changes of the system height. An attractive feature of the compressional beta effect is that it does not require large-scale coherent flow structures that touch the spherical boundaries in order to operate. Instead it entirely relies on the local expansion and compression of the fluid. Indeed, \citet{Glatzmaier2009} argue that the topographic beta effect is hard to maintain in the presence of the vigorous turbulence expected in planetary interiors. In their view, coherent fluid columns connecting the boundaries are likely to be shred apart by the vigorous convection, especially in planetary interiors which exhibit a strong density stratification which further facilitates the break-up of convection columns. The slope of the remote boundaries thus seems unlikely to efficiently generate vorticity in the deep interior, which, as argued by \citet{Glatzmaier2009}, casts doubt on the relevance of the {Busse-} and the {topographic Rhines picture} for highly turbulent, strongly density stratified planetary atmospheres. Models based on the compressional beta effect offer an interesting alternative here. Recently, \citet{Evonuk2006,Evonuk2007,Evonuk2008,Glatzmaier2009} and \citet{Evonuk2012} put forward a model that shares many similarities with the Busse picture, but relies on the compressional beta effect instead of the topographic one. Structure models of giant planets reveal that the density scale height in giant planets changes considerably with depth. Since the phase velocity of the Rossby waves generated by the compressional beta effect depends on the local density scale height, the waves must be expected to propagate with different phase speed at different depth, which, as in the Busse picture, would tilt convective plumes. For Jupiter- or Saturn-like density profiles, Rossby waves propagate faster near the surface than in the deep interior. Again similar to the Busse picture, the arising Reynolds stresses then transport prograde momentum outwards and retrograde momentum inwards, a process that is subsequently reinforced by a mean flow instability. This naturally explains the strong prograde equatorial jets of Jupiter and Saturn. The retrograde jets found on Uranus and Neptune may also be explained by density profiles with radially increasing density scale heights close to the surface, as suggested by models of their interior structure \citep{Hubbard1991}. These ideas have been tested using two-dimensional numerical simulations in equatorial planes using the anelastic approximation \citep{Evonuk2006,Evonuk2007,Evonuk2008,Glatzmaier2009,Evonuk2012}. The simulations with an inner core typically reveal the formation of two jets, one touching the planetary surface at the equator, and another one of opposite direction closer to the rotation axis, with the jet directions being controlled by the background density profile as expected. Moreover, fully convective models without an inner core allow for the occurrence of three jets. The authors of these studies provide a detailed description of the mechanistic processes generating these jets. We refer to this view of zonal flow generation by the term {\em Evonuk / Glatzmaier picture} in table \ref{mechanism_table}. Interestingly, \citet{Evonuk2012} briefly mention that for rapid rotation and weak convective forcing, the simulations show a tendency to develop more jets, without further discussing such multiple jet states. Taking into account the similarities between the Evonuk / Glatzmaier model with the Busse picture, and the related nature of the corresponding Rossby waves, it seems likely that the topographic Rhines picture should also have a compressional equivalent. As already suggested in the early works of \citet{Ingersoll1982} and \citet{Ingersoll1986}, who mainly focussed on the stability properties of barotropic jets, such a {\em compressional Rhines mechanism} may then be expected to drive multiple zonal flows. In table \ref{mechanism_table}, we call this view of zonal flow generation the {\em compressional Rhines picture}. A detailed description will be provided in section \ref{theory}. So far, convincing experimental or numerical validation of this concept is still missing. In this paper, we aim to close this gap. The numerical evidence presented in the following sections suggests that the compressional Rhines picture provides an interesting avenue for combining the attractive features of the Evonuk / Glatzmaier view with the ability to explain multiple jets in a straight-forward manner. A number of questions immediately arise in this context. Most importantly, does a compressional Rhines mechanism indeed generate multiple jets in rapidly rotating, convective turbulence? The answer is far from obvious, because in convective systems, the strength and spatial distribution of the turbulent forcing is intimately tied to the large scale dynamics. The natural tendency of convective turbulence to radially mix the fluid homogeneously has to be overcome in order to allow for a radially inhomogenous potential vorticity (PV) distribution, as required for jets. In particular, it appears unclear whether eddy-transport barriers, as envisioned e.g. by \citet{Dritschel2008}, can be maintained against incursions of vigorous convective plumes. If jets can indeed be generated in the proposed way: How well does the usual picture of turbulence-wave crossover describe the dynamics in direct numerical simulations? Do the jets generated by a compressional Rhines mechanism follow a compressional Rhines scaling? What are the numerical parameters characterizing such a scaling law? Do predictions based on this scaling broadly fit the planetary jets observed in our solar system? Preliminary answers to these question, based on a highly idealized \jan{2-d} model of planetary convection, will be given in the following sections. Our paper is organized as follows. To give the reader a visual idea of the physical effects discussed in this work, we first describe some observations from numerical simulations of compressible (anelastic) rotating convective turbulence developing into multiple jets (section \ref{observation}). A more formal definition of the model considered in this paper is given in section \ref{method}, which is then used to develop a simple theoretical model for the observed jets in section \ref{theory}. The theoretical predictions are compared with the results from a large number of numerical simulations in section \ref{results}. Following some speculations on the implications of our work for planetary interiors in section \ref{speculations}, general conclusions are given in section \ref{conclusions}. | \label{conclusions} \stephan{As described in the introduction, Rhines-type \neu{dynamics}\alt{$\beta$-mechanisms} \alt{are} \neu{is} a recurring theme in theories of planetary jet dynamics, occurring in latitudinal, topographic and compressional form, as \alt{is their competition with} \neu{are} other \neu{mechanistic pictures}\alt{mechanisms} \citep[e.g.][]{Busse2002,Glatzmaier2009} \alt{relying on} \neu{that emphasize} spatial variations in the strength of the beta term. While the \stephannew{latitudinal} and the topographic beta effects have been studied extensively in the literature, in this paper we focussed on a \alt{compressional} \neu{compressional Rhines-type scenario}\alt{$\beta$-mechanism}, which has received little attention since it was originally proposed by \citet{Ingersoll1982}. Especially in the context of the jet generation mechanism recently introduced and analyzed by \citet{Evonuk2006,Evonuk2007,Evonuk2008,Glatzmaier2009} and \citet{Evonuk2012}, \neu{this} \alt{compressible $\beta$-mechanism} appears to be an interesting alternative \neu{view}. } \stephannew{Indeed,} \stephan{our study clearly demonstrates that rapidly rotating, turbulent convection has the ability to drive a multitude of deep jets by a\alt{ Rhines-type} compressional \neu{Rhines-type} \alt{$\beta$-}mechanism. Results from a suite of two-dimensional numerical simulations employing the anelastic approximation reveal that -- as theoretically expected -- the typical jet width is well predicted by a compressional Rhines scale $l_R$, which depends on the planetary rotation rate, on the \neu{typical density scale height}\alt{amount of density increase with depth}, and on the \neu{jet velocity}\alt{vigor of the flow}. Kinetic energy spectra clearly show the anticipated turbulence-wave crossover. \neu{Furthermore,} potential vorticity staircases are found to develop, which are accompanied by corresponding, unstably stratified staircases in the \stephannew{potential density field. While the steplike distribution of potential vorticity appears to suppress the turbulent mixing across the interface regions, the buoyancy jumps over these interfaces act to facilitate convective transport.} Such counteracting mechanisms are absent in homogenous beta-plane turbulence driven by a prescribed forcing \neu{(e.g. \citet{Vallis1993,Rhines1994,Vallis2006,Scott2012a,Scott2012b})}, which shows that the robustness of the compressional \neu{Rhines picture}\alt{$\beta$-mechanism} found in this paper is not \neu{obvious from the outset}\alt{a-priorly obvious}. The richness of the dynamics is further illustrated by the occurrence of relaxation oscillations, resulting from the interaction between the convection and the jet shear. \neu{Similar behavior has been reported in previous studies of convection in the presence of a topographic beta effect (extensive references are given in section \ref{observation}). } \alt{In all cases considered, the compressional $\beta$-mechanism dominates over the Evonuk / Glatzmaier mechanism as soon as the system becomes deep enough to host multiple compressional Rhines lengths $l_R$. It thus naturally explains the occurrence of multiple jets, and operates even for a depth-independent density scale height, a condition under which the Evonuk / Glatzmaier mechanism breaks down. In conclusion, we expect that the compressional-beta mechanism dominates in systems deeper than $O(l_R)$, while the Evonuk / Glatzmaier mechanism takes over otherwise. Together, both pictures provide a coherent picture of the role of compressibility in generating zonal winds.} \stephan{Whether\alt{ or not} the compressional \neu{beta effect}\alt{$\beta$-mechanism} can be expected to \neu{drive deep multiple jet patterns} \alt{contribute to deep jet generation } in planetary objects \alt{thus seems to depend} \neu{depends} on the magnitude of the compressional Rhines lengths. A naive application to giant planets indicates that it should indeed be important in Jupiter and Saturn, where the scaling law found in this study even results in reasonable predictions for the observed number of surface jets. Somewhat more unexpected is the fact that in the absence of magnetic fields, a large number of jets is predicted \neu{also} for the Earth's outer core. Although such jets are expected to be damped by Lorentz forces, this result illustrates that the Boussinesq models typically used in this context might miss important physical effects.} \stephan{The highly speculative nature of such applications to planetary bodies should be pointed out explicitly here. The simple model investigated in this paper was designed with the goal to demonstrate the efficiency of the compressional \neu{Rhines-type dynamics}\alt{beta mechanism} in a convecting system as simply and clearly as possible. Physical processes which are not essential for \neu{such a mechanism}\alt{this mechanism} have been intentionally neglected, irrespective of their possible importance in real systems, with the goal to study the compressional beta effect in isolation. } \stephan{Perhaps the most important limitation of our model is that it is restricted to two-dimensional flows in an equatorial plane. It may be argued that nearly two-dimensional flows may indeed be expected there because rapid rotation tends to largely suppress flow variations along the rotation axis \neu{(e.g. \citet{Busse2002,Schaeffer2005,Calkins2012})}. However, the question to which degree processes \neu{only present in three dimensions} like vortex stretching and tilting can be safely neglected, and whether an inverse cascade will indeed occur, need further investigation. Encouraging in this context is the identification of an inverse cascade in numerical simulations of three-dimensional rapidly rotating Rayleigh-B\'enard convection using a reduced, asymptotic model \citep{Julien2012}. Different from the classical and the topographic beta effect, which produce vorticity for fluid volumes moving latitudinally or perpendicular to the rotation axis, the compressional beta effect needs radially moving fluid parcels to operate. The consequences for the dynamics in three-dimensional spherical shells have not been addressed in this paper and need further attention. } \stephan{Another open issue is the competition with the other mechanisms proposed in the literature. }\stephannew{\neu{Processes}\alt{Effects} like the topographic or \neu{classical}\alt{latitudinal} \neu{beta effects}\alt{$\beta$-mechanisms} are absent from our model by construction.} \stephan{Simulations in three-dimensional spherical shells are desirable in this context. Our study however suggests that large rotation rates, corresponding to Ekman numbers below $ O(10^{-7})$, \stephannew{are} required to observe a compressional \neu{Rhines }\alt{$\beta$-}mechanism \jan{in low-latitude regions} for $O(1)$ Prandtl number fluids}. \stephan{Such Ekman numbers are currently not reached in numerical simulations of convection in rotating spherical shells. Among such models, only a few have included the effect of a density increase with depth so far \citep{Kaspi2009,Jones2009,Gastine2012}, typically using Ekman numbers between $O(10^{-4})$ and $O(10^{-6})$.} \neu{These highly resolved 3-d simulations as well as their incompressible counterparts (e.g. \citet{Heimpel2005,Heimpel2007}) consistently only find two jets on the equatorial plane.} \stephannew{An interesting first step towards identifying a compressional \neu{Rhines }\alt{$\beta$-}mechanism in rotating spherical shells has recently been carried out in parallel to our study by \citet{Gastine2014}. Although their work is restricted to moderate Ekman numbers, the authors argue that the typical length scales for high latitude jets are better described by a compressional Rhines scale than by its topographic counterpart. Both effects are however difficult to separate, because the expected scaling difference is small in the accessible parameter range. The next generation of numerical convection models will hopefully allow us to enter the relevant low Ekman number regime, such that the relative strength and mutual interaction of the compressional beta effect and other relevant mechanisms can be studied in detail. Until then, the study presented in this paper may serve at least as clear and unambiguous evidence for the general ability of \neu{the}\alt{a Rhines-type,} compressional \neu{beta effect}\alt{$\beta$-mechanism} to \neu{generate}\alt{drive} multiple zonal jets in convective systems.} \jan{Is the compressional beta effect indeed a source of planetary zonal winds? Although we cannot answer this question here, the main conclusion of our study is that it should be seen as one of several possible candidates involved in driving (deep) planetary jet patterns. } | 14 | 4 | 1404.6940 |
1404 | 1404.3562_arXiv.txt | With application to inner stellar radiative zones, a linear theory is used to analyze the instability of a dipole-parity toroidal background field, in the presence of density stratification, differential rotation, and realistically small Prandtl numbers. The physical parameters are the normalized latitudinal shear $a$ and the normalized field amplitude $b$. Only the solutions for the wavelengths with the maximal growth rates are considered. If these scales are combined to the radial values of velocity, one finds that the (very small) radial velocity only depends slightly on $a$ and $b$, so that it can be used as the free parameter of the eigenvalue system. The resulting instability-generated tensors of magnetic diffusivity and eddy viscosity are highly anisotropic. The eddy diffusivity in latitudinal direction exceeds the eddy diffusivity in radial direction by orders of magnitude. Its latitudinal profile shows a strong concentration toward the poles which is also true for the effective viscosity which has been calculated via the angular momentum transport of the instability pattern. The resulting effective magnetic Prandtl number reaches values of $O(10^2)$, so that the differential rotation decays much faster than the toroidal background field, which is {the} necessary condition to explain the observed slow rotation of the early red-giant and sub-giant cores by means of magnetic instabilities. | Hydromagnetic dynamos can be understood as a magnetic instability driven by a flow pattern in fluid conductors. There are, however, strong restrictions on the characteristics of flows that can excite dynamos \citep{E46}, as well as on the geometry of the resulting magnetic fields \citep{C33}. The restrictions even exclude any dynamo activity for a number of flows, for example, differential rotation alone can never maintain a dynamo. An open question is whether magnetic instabilities are able to excite a sufficiently complicated motion that together with the background flow can generate magnetic fields. \citet{TP92} suggested that nonuniformly rotating disks can produce a dynamo when magnetorotational (MRI) and magnetic buoyancy instabilities are active. Later on, numerical simulations have shown that the MRI alone may be sufficient for the accretion disk dynamo. However, at least for low magnetic Prandtl numbers, it remains unclear whether the MRI-dynamo is physical, or just a numerical artifact. Another possibility was discussed by Spruit (2002), who suggested that differential rotation and a magnetic kink-type instability \citep{T57,T73} can together drive a dynamo in stellar radiation zones. If real, such a dynamo would be very important for angular momentum transport in stars. It taps energy from differential rotation, thus reducing the rotational shear. Radial velocity fluctuations converting toroidal magnetic field into poloidal field are necessary for such a dynamo. This dynamo, therefore, would necessarily also mix chemicals in stellar interiors, which may have consequences for stellar evolution. However, simply the action of differential rotation and a magnetic instability converting toroidal field into poloidal does not guarantee such a dynamo, which has thus far not been proven to exist. Doubts are especially directed to the kink-type instability which, in contrast to the MRI, exists even without differential rotation and develops at the expense of the magnetic energy. Detailed estimations of the dynamo parameters are thus necessary to assess any dynamo-efficiency of the Tayler instability. A basic aspect of turbulent dynamos is the ability of correlated magnetic (${\vec b}$) and velocity (${\vec u}$) fluctuations to produce a mean electromotive force along the background magnetic field ${\vec B}$, and also along the electric current $\vec J$, i.e. \begin{equation} \langle {\vec u}\times{\vec b}\rangle = \alpha{\vec B}- \mu_0 \eta_{\rm T}{\vec J} . \label{1} \end{equation} We computed the $\alpha$ effect of the current-driven instability for rigid rotation and for a 3D magnetic geometry in a previous paper (R\"udiger et al. 2012). However, because a magnetic-instability-induced radiative-zone dynamo requires a differential rotation as an energy source, it remains to find the instability and the resulting electromotive force in the presence of a (weak) differential rotation. We have shown earlier that the interaction of even current-free toroidal fields and differential rotation leads to the appearance of an own instability where the growth rates (for fixed magnetic field amplitude) switch from the low diffusion frequency to the high rotation frequency. It would therefore not be surprising if the current-driven instability under the presence of differential rotation exceeded the power of the instability with no rotation or with rigid rotation. Indeed, for toroidal fields in cylindrical geometry one can show that for rigid rotation the instability is suppressed for too weak toroidal magnetic field (Acheson 1978; Pitts \& Tayler 1985; Tataronis \& Mond 1987), while it is re-animated by differential rotation (see R\"udiger et al. 2013). In the present paper we shall apply a linear theory to show that similar results are obtained even in non-uniformly rotating spheres with northern and southern magnetic belts of opposite direction. The formation of red giants without internal angular momentum transport would lead to steep radial profiles of the internal rotation law. Ceillier et al. (2012) report the Kepler-like profile $\Om \propto r^{-1.6}$ for the low-mass red giant KIC 7341231. The {\em Kepler} data, however, speak another language. According to the asteroseismology results, the core rotates only slightly faster than the outer convection zone (Deheuvels et al. 2012, 2014). Eggenberger et al. (2012) argue that an artificial viscosity of $3 \times 10^4$ cm$^2$/s may explain the surprisingly flat rotation profiles. The outward flux of angular momentum then produces the observed spin-down of the inner radiative core. The general condition for this mechanism to work is that {\em the rotation profile decays faster than the magnetic field}. The magnetic Prandtl number represents the ratio of the decay time of the magnetic field and the decay time of the differential rotation. This basic requirement is thus fulfilled when the effective magnetic Prandtl number is much larger than unity. We have thus to study whether and under what circumstances the magnetic Prandtl number becomes large. The model and the stability analysis of this paper are close to those of \citet{KR08} and will be discussed here only briefly. The main idea is the use of the very small but necessarily nonzero radial flow component as the free parameter of the eigenvalue system, which is assumed to be proportional to the product of the critical length scale and the growth rate of the unstable mode. The special choice of this parameter always becomes unimportant for the results, which can be understood as the ratio of second-order quantities as there are the energy ratios and/or the effective magnetic Prandtl number formed by the diffusivity coefficients. The dominant component of the magnetic field inside a star is assumed to be the toroidal one, which can easily be produced by differential rotation by induction from a weak poloidal field. The background toroidal field of our model consists of two latitudinal belts of opposite polarities \begin{equation} {\vec B} = r\sin\theta\sqrt{\mu_0\rho}\ \Om_\mathrm{A}(\mu) {\vec e}_\phi \label{2} \end{equation} (see Spruit 1999) with $\Om_\mathrm{A}(\mu)$ as the Alfv\'en frequency of the toroidal field. Spherical coordinates are used with the axis of rotation as the polar axis; ${\vec e}_\phi$ is the azimuthal unit vector and $\mu= \cos\theta$. Field belts with equatorial symmetry are described by functions $\Om_\mathrm{A}$ which are symmetric with respect to the equator. On the other hand, the simplest case with two belts of opposite signs in the two hemispheres is $\Om_\mathrm{A}= b\ \Om\ \mu$. The latitudinal profile of (\ref{2}) then peaks in mid-latitudes at $\theta=45^\circ$ and $\theta=135^\circ$. The quantity $b$ is referred in the following as the {\em amplitude} of $\hat\Om_\mathrm{A}=\Om_\mathrm{A}/\Om$. Within the solar tachocline toroidal fields of (say) 1 kG are represented by $b\simeq 0.01$. The background flow is simply \begin{equation} {\vec U} = r\sin\theta \ \Om \ {\vec e}_\phi \label{3} \end{equation} with $\Om$ as the equatorial rotation rate. For $\Om_\mathrm{A}=\Om={\rm const}$, i.e. $b=1$, one finds $ {\vec U}={\vec V}_{\rm A}$ with ${\vec V}_{\rm A}$ the Alfv\'en velocity, which is stable according to the formulation of Chandrasekhar (1956) for ideal, perfectly-conducting fluids. More realistic $\Om_\mathrm{A}$ and $\Om$ are radius-dependent, but this dependence is not of basic importance for the analysis of barytropic instabilities. The reason is that the stratification of the radiative core is stable with positive \begin{equation} N^2 = \frac{g}{C_\mathrm{p}}\frac{\partial S}{\partial r} , \label{4} \end{equation} where $S$ is the specific entropy and $C_\mathrm{p}$ the specific heat for constant pressure. We shall see that the latitudinal shear of the global rotation is much more important. Let the rotation law be given by \begin{equation} \Om= \Om_0 (1 - a \ \cos^2\theta) \label{Om} \end{equation} with $a$ as the shear parameter of the normalized differential rotation. $\Om_0$ is the angular velocity of the equator, positive $a$ describes (`solar-type') equatorial acceleration. This is the typical definition for the tachocline theory: the outer convection zone above the radiative interior generates equatorial acceleration at its bottom which defines the shear at the top of the radiation zone. | The instability of a large-scale toroidal magnetic field system with dipolar parity under the influence of a global rotation with latitudinal shear has been studied. We are able to compute simultaneously the mean stress tensor and the mean electromotive force from one and the same model, which allows the determination of both the effective diffusion coefficients $\eta_{\theta\theta}$ and $\nu_{\rm H}$. The solution of such linear equation systems always contains an unknown arbitrary amplitude factor, which we have fixed in the present paper by the relation $u_r=\alpha \lambda \omega_{\rm gr}$ for the radial flow of the instability, with $\lambda$ as the critical wavelength and $\omega_{\rm gr}$ as the maximal growth rate of the modes. The factor $\alpha$ plays the role of the free parameter of the system. Almost all of the numerical applications in this paper concern the upper layer of the solar radiative zone, also called the solar tachocline. The buoyancy frequency in this region exceeds the rotation rate by a factor of about 400, which leads to rather slow radial perturbations in comparison to the horizontal ones. If the Alfv\'en frequency of the toroidal field is taken in units of the rotation rate, i.e. $b\simeq\Om_{\rm A}/\Om$, then weak fields are described by $b\simeq 0.01$ and strong fields by $b\simeq 1$. For rigid rotation and for $\alpha\simeq 1$ the resulting radial flow for the instability with $b\gsim 0.5$ is of order 1 mm/s. The simplest estimate for the diffusion coefficient $\eta_{\rm T}\simeq \lambda^2\omega_{\rm gr}$ peaks for $b\simeq 0.1$ and reaches values of 250 in units of the microscopic diffusivity. This value would lead to a decay of the toroidal magnetic background field after 0.5 Gyr. If the differential rotation shall decay within this time the condition $\rm Pm_{\rm T}>1$ must be fulfilled, otherwise the magnetic field decays faster than the differential rotation. It is not possible for MHD flows to estimate the effective magnetic Prandtl number by simple scale relations, because Reynolds and Maxwell stresses contribute differently to the two diffusivities. To determine both the viscosity and the magnetic diffusivity from one and the same model one has to solve the full equation system and to form the correlations which define the two diffusivities. In ratios of expressions of second order such as the electromotive force and/or the Reynolds/Maxwell stress the free parameter $\alpha$ no longer influences the results. The inspection of the correlations which provide eddy diffusivity and eddy viscosity reveals an extreme anisotropy of both diffusivity tensors. As the radial components of flow and field of the instability pattern are smaller by orders of magnitudes than the horizontal components, the correlations including $u_r$ and $b_r$ are also smaller by orders of magnitudes than the horizontal correlations. One must thus be careful with the definition of the magnetic Prandtl number. In this paper only the tensor components of the diffusivities are considered which dissipate both the latitudinal gradients of the background flow and field. Hence, the expressions (\ref{ub}) and (\ref{vis1}) must be computed. The relation (\ref{Pm}) defines the magnetic Prandtl number as the ratio of the diffusion coefficients which result from the Eqs. (\ref{etar}) and (\ref{vis2}). We indeed find values up to $O(10^2)$ for the effective Prandtl number at $b\simeq 0.1$. The magnetic diffusivity, however, grows monotonically for growing $b$. Strong fields with $b>1$ thus decay very quickly. For weaker background fields ($b\simeq 0.1$) the latitudinal shear of the global rotation decays 100 times faster than the field itself. One might thus predict that already after (say) 1 Myr the radiative stellar cores rotate as a solid body. This, however, is not completely true. The calculations show that the magnetic instability transports angular momentum equatorward even for solid-body rotation. This so-called $\Lambda$ effect avoids the formation of uniform rotation in the same way as it does in rotating solar and stellar convection zones where it is directly observable. Rigid rotation is no longer a solution of the stationary Reynolds equation with Lorentz force. As the density gradients and the horizontal anisotropies are weaker the $\Lambda$ effect for magnetized radiative zones, however, is much smaller than in rotating convection zones. Nevertheless, also weak differential rotation permanently produces toroidal fields from fossil poloidal fields, so that their existence during the existence of magnetic instabilities must be the rule rather than the exception. | 14 | 4 | 1404.3562 |
1404 | 1404.1084_arXiv.txt | In the solar system, moons largely exceed planets in number. The {\it Kepler} database has been shown to be sensitive to exomoon detection down to the mass of Mars, but the first search has been unsuccessful. Here, we use a particles-in-cell code to predict the transit of the plasma torus produced by a satellite. Despite the small size of a moon, the spatial extent of its plasma torus can be large enough to produce substantial transit absorptions. The model is used for the interpretation of {\it Hubble Space Telescope} early ingress absorptions apparently observed during WASP-12\,b and \hdn\,b UV transits for which no consistent explanation exists. For \hdn\,b an exomoon transiting $\sim 16$\,$R_p$ ahead of the planet and loading $\sim 10^{29}$ C\,II ions/s into space is required to explain the tentative early ingress absorption observed for C\,II. For WASP-12b, a moon transiting $\sim 6$\,$R_p$ ahead from the planet and ejecting $\sim 10^{28}$ Mg\,II ions per second is required to explain the NUV early ingress absorption feature. Interestingly, both \hdn b and WASP-12b predicted satellites are outside the Hill sphere of their planets, an indication that the moons, if present, were not formed in situ but probably captured later. Finally, our simulations show a strong electromagnetic coupling between the polar regions of planets and the orbital position of the moons, an expected outcome of the unipolar induction DC circuit model. Future observations should test our predictions with a potential opportunity to unambiguously detect the first exomoon plasma torus. | Moons represent a fundamental component of our solar system because they are tightly related to its formation and evolution process. Regular moons represent a sub-population that was formed in situ and evolved with their parent planets. A second sub-population could have been captured after the formation of the planet, leading to a system with an alien body that may have a distinct composition \citep{hua83}. As seen from the planet, a group of satellites represents a complex gravitational system that has several dynamical properties comparable to a planetary system. Therefore, detecting satellites can enhance our understanding of planetary system evolution by extending the spectrum of gravitational configurations of {\it N}-body systems. \begin{figure} \centering \includegraphics[width=6cm,angle=90]{f1.eps} \caption{{ Three-dimentional PIC simulation of \hdn\,b (or WASP-12b) magnetosphere under sub-Alfv\'enic stellar flow conditions \citep{coh11}. The planet position (0,0) is indicated by a small blue disk. The magnetopause position is also shown. The simulation was obtained for a stellar wind that has a relative speed of $230$\,km/s, an Alfv\'en speed of $\sim 300$\,km/s corresponding to an interplanetary magnetic field of $\sim 4$\,mG, and a sound speed of $\sim 190$\,km/s leading to a sonic Mach number $M_s\sim 1.2$ and Alfv\'en Mach number of $M_v\sim 0.8$. Because C\,II and Mg\,II abundances are negligible in the stellar wind, this structure has no signature in the corresponding transit light curves. Other relevant parameters are discussed in the text.} } \end{figure} The large number of satellites in our solar system and the promise of the increasing number of exoplanets being discovered adds to the likelihood of detection of an extrasolar satellite in the near future. Several techniques have been proposed to detect exomoons principally based on the transit technique \citep{sar99,kip11,sat09,sim12}. Unfortunately, both the small size and mass of moons make their detection a very difficult task \citep{kip11}. Apart from technological limitations, one potential explanation for the non-detection of exomoons could be that for most detected exoplanets orbiting very close to their host stars, the size of their Hill sphere is shrunk significantly, probably limiting their satellite population mostly to bodies captured after the exoplanet formation \citep{dom06,nam10,wei10}. \begin{figure*} \centering \hspace*{-0.4in} \includegraphics[width=9cm]{f2a.eps} \includegraphics[width=9cm]{f2b.eps} \hspace*{-0.4in} \includegraphics[width=9cm]{f2c.eps} \includegraphics[width=9cm]{f2d.eps} \caption{Three-dimentional PIC simulations of exomoon plasma tori evolving within the magnetosphere cavity shown in Figure 1. Satellite orbits shown in yellow. Top left: plasma distribution of Mg\,II ions in the orbital plane for the case of WASP-12b. Extrasolar satellite orbiting at $\sim 6$\,$R_p$ from its parent planet and losing Mg\,II ion plasma at a rate of $\sim 5\times 10^5$\,g/s. Satellite at 9 o'clock. Top right: same but in the meridian plane. Bottom left: plasma distribution of C\,II ions in the orbital plane for the case of \hdn\,b. Satellite orbiting at $\sim 16$\,R$_p$ from its parent planet and losing C\,II ion plasma at a rate of $\sim 2\times10^6$\,g/s. Satellite at 3 o'clock. Bottom right: same as in left but in meridian plane.} \end{figure*} In the description adopted in most of the existing literature regarding exoplanetary environments, one key component has been rarely cited: the occurrence of a plasma torus that may form around an exoplanet by an orbiting moon, much like the Io torus around Jupiter \citep{kup76,bro76}. Emission lines from hot plasma tori have been indeed proposed to uncover extrasolar satellites, yet no quantitative diagnostic has been put forth \citep{mar00}. Here, we investigate the transit absorption by a plasma torus formed by a satellite orbiting an exoplanet using a particle-in-cell (PIC) code \citep{ben13}. After a short description of our PIC model and the magnetospheric configurations of \hdn b and WASP-12b that are considered here as illustrations of the proposed new diagnostics, we will show that transit features such an early ingress absorption may appear, depending on the mass loss rate and the orbital parameters of the assumed moon. | Here, we propose using plasma tori that are produced by volcanic moons along their orbits to detect small-size and small-mass satellites around exoplanets. Following the Io torus example, a small body with enough volcanic activity may produce a spatially extended plasma nebula that may show substantial transit absorption, particularly in the UV. The proximity to a parent star should likely enhance the volcanic activity of the small body, inherently enhancing the mass loading the planet's magnetosphere. It is therefore reasonable to posit that a strong transit signature might then be detected. To test this scenario, we utilized our PIC 3D simulation code in which we placed a moon in orbit inside the magnetosphere cavity of a magnetized exoplanet. Without any loss of generality, we considered {\it specific} plasma conditions of exoplanet \hdn\,b and WASP-12b for illustration. For the selected cases, we show clearly that the transit absorption of a torus may be strong and varying from ingress to egress. In the example shown in Figure 4 for exoplanet \hdn\,b and C\,II line conditions, a strong early ingress absorption may appear for specific orbital parameters of the moon when assuming a $\sim 2\times 10^{6}$\,g/s outflow of C\,II ions from the small body. The total mass loss depends on the dominant species (mean mass) and the relative abundances, yet it should not exceed $\sim 2.2\times 10^{16}$\,g/year or $\sim 2\times10^{-10}$ Io$_{\rm mass}$/year. In the case of WASP-12b, a good fit to the observed Mg\,II transit light curve was obtained with a moon mass loss of $\sim 5 \times 10^{5}$\,g/s of Mg\,II ions (e.g., Fig. 3). Depending on the moon composition, the total mass loss should not exceed $\sim 1.8\times 10^{16}$\,g/year, a rate comparable to the \hdn\, value found above. Both exomoons orbital properties are consistent with the periodic and stable family {\it f} of the restricted three-body problem \citep{hen70}. With all the predictions proposed here, new transit observations should be obtained to definitely confirm the early ingress signature, particularly for \hdn , a bright target for which isolated ion resonances lines (like C\,II) allow a relatively easy Doppler analysis of the absorption. | 14 | 4 | 1404.1084 |
1404 | 1404.7561_arXiv.txt | Spectral measurements of the $21$\,cm monopole background have the promise of revealing the bulk energetic properties and ionization state of our universe from $z\sim 6-30$. Synchrotron foregrounds are orders of magnitude larger than the cosmological signal, and are the principal challenge faced by these experiments. While synchrotron radiation is thought to be spectrally smooth and described by relatively few degrees of freedom, the instrumental response to bright foregrounds may be much more complex. To deal with such complexities, we develop an approach that discovers contaminated spectral modes using spatial fluctuations of the measured data. This approach exploits the fact that foregrounds vary across the sky while the signal does not. The discovered modes are projected out of each line of sight of a data cube. An angular weighting then optimizes the cosmological signal amplitude estimate by giving preference to lower-noise regions. Using this method, we show that it is essential for the passband to be stable to at least $\sim 10^{-4}$. In contrast, the constraints on the spectral smoothness of the absolute calibration are mainly aesthetic if one is able to take advantage of spatial information. To the extent it is understood, controlling polarization to intensity leakage at the $\sim 10^{-2}$ level will also be essential to rejecting Faraday rotation of the polarized synchrotron emission. | One of the richest yet least understood narratives in cosmology is the formation of the complex structure that we see today out of the simple initial conditions implied by the cosmic microwave background (CMB). The first luminous objects are thought to have formed at $z\sim 20-30$ through collapse in $10^6-10^8\,{\rm M}_\odot$ halos \citep{2001PhR...349..125B, 2013RPPh...76k2901B}. The radiation from these objects heated and then reionized the intergalactic medium (IGM). There are several sources of complementary information about the evolution of ionization in this epoch. The CMB temperature anisotropy is damped by the total Thomson depth to free electrons. The Planck collaboration has used this effect, combined with a constraint on the scalar amplitude from gravitational lensing, to measure the optical depth through reionization \citep{2013arXiv1303.5076P}. In addition, Thomson scattering through the reionization epoch generates a unique polarization signature on large angular scales. WMAP has measured the total optical depth using this polarization signature \citep{2013ApJS..208...20B}. These are integral constraints on the free electron abundance and can be translated into a central reionization redshift of $10.6 \pm 1.1$ \citep{2013ApJS..208...20B}. Once the IGM is highly ionized, it is transparent to Ly-$\alpha$ photons. Absorption measurements along sight lines to high-redshift quasars indicate that reionization must have ended by $z<6$ \citep{2006AJ....132..117F}. Absorption saturates at low abundance, so these should be taken as bounds on the end of reionization, which could still have been largely complete at redshifts higher than $6$. Recently, two methods have been developed to place much more direct bounds on the duration of reionization. The ionization process is thought to be spatially patchy, as local sources of radiation blow ionized bubbles that coalesce into the fully reionized IGM. CMB photons scattering from this patchy screen produce an additional kinetic Sunyaev-Zel'dovich anisotropy appearing most clearly at $\ell > 3000$, where the primary CMB is negligible \citep{1998PhRvL..81.2004K}. Upper limits on this effect translate into a model-dependent upper bound on the duration of reionization \cite{2012ApJ...756...65Z} and hold the promise of direct detection of patchy structure in the near future. The patchy structure of reionization can also be observed directly in three dimensions using emission of neutral hydrogen through its $21$\,cm line \citep{2006PhR...433..181F}. Recent bounds from GMRT \citep{2013MNRAS.433..639P}, MWA \citep{2014PhRvD..89b3002D}, and PAPER \citep{2014ApJ...788..106P, 2013ApJ...768L..36P} are marching down to the expected level of fluctuations, in parallel to efforts at LOFAR \citep{2013A&A...556A...2V}. An alternative to measuring the $21$\,cm anisotropy is to measure the signal of its global emission (or absorption at earlier times) \citep{2008PhRvD..78j3511P, 2010PhRvD..82b3006P}, which reveals the bulk energetic properties and ionization state of the universe during reionization and preceding epochs when the first luminous structures were forming. Global $21$\,cm experiments include EDGES \citep{2010Natur.468..796B} and SCI-HI \cite{2014ApJ...782L...9V} (which have both reported bounds), LEDA\footnote{{\tt http://www.cfa.harvard.edu/LEDA/science.html}} and the proposed DARE mission \citep{2012AdSpR..49..433B}. The frequencies of interest in these global studies are $\sim 50-200$\,MHz, and fiducial theoretical models suggest a maximum contrast of $\sim 100$\,mK relative to the synchrotron emission of the galaxy, which can vary $\sim 10^2-10^5$~K across the sky and frequency range. Astrophysical synchrotron emission is thought to be fully described by a handful of spectrally smooth functions that can be distinguished from the variation of the global reionization signal. Extremely bright foregrounds make the measurement susceptible to instrumental systematics. For example, if an instrument with a $1\%$ perturbation to the spectral calibration observed a $500$~K power law, subtraction of a pure power law would leave a $5$~K residual, significantly larger than the signal. Through careful instrumental design, the level of instrumental systematics may be controlled but generally cannot be nulled entirely. Here, we develop methods that can be used to constrain the cosmological signal in these heavily contaminated data. Following the monopole nature of the signal, experiments to date have mapped the sky with very large beams \citep{2010Natur.468..796B, 2014ApJ...782L...9V}. However, a unique trait of the foregrounds is that they vary across the sky, while the signal is constant. This regime is the opposite of the situation normally found in analysis of small signals, where a modulated signal is pulled out of foregrounds. \citet{2013PhRvD..87d3002L} (henceforth L13) proposed that experiments seeking to measure the global signal should also resolve the sky with an instrumental beam. This allows selective weighting against regions of high contamination and allows for the use of angular correlation information to reject foregrounds. The additional spatial resolution yields higher fidelity recovery of the cosmological $21$\,cm spectrum. Here, we extend the ideas in L13 to a method that uses the spatial fluctuations of foregrounds in the data to discover contaminated spectral modes. A similar idea has been employed successfully in $21$\,cm large-scale structure measurements \cite{2013MNRAS.434L..46S, 2013ApJ...763L..20M, 2010Natur.466..463C} at $z\sim 1$ and has been suggested for cleaning ionospheric contamination \citep{2014MNRAS.437.1056V}. Discovery of foreground spectral modes in the measured data makes the method more robust to assumptions about the foreground contamination. For example, now if the instrumental passband has a $1\%$ ripple, the largest foreground mode discovered in our foreground cleaning method will also self-consistently exhibit this ripple. Generally, instrumental systematics take relatively clean and smooth functions of frequency from synchrotron emission and convert them into a more complex structure that requires additional spectral functions to describe. We argue that the primary goal in instrumental design should be to prevent proliferation of bright, new foreground modes in the data. Each new foreground degree of freedom produced by instrumental response to foregrounds results in more signal loss and makes discovery of the signal more ambiguous. The methods described here of (1) using spatial variation to discover spectral foreground modes, which can then be projected out and (2) down-weighting known spatial areas of high contamination (the galaxy) provide the strongest methods for recovering the global $21$\,cm signal in the absence of additional prior information about the foregrounds or instrumental response. While the algorithm of mode subtraction and angular weighting is intuitive, we develop it from the ground up to expose several implicit choices and possible pitfalls. Recently, \citet{2014arXiv1404.0887B} argued that a dipole gain pattern can be calibrated in an interferometric array. However, additional variations in spectral response due to factors such as the analog-to-digital converter, reflection, and signal loss after the antenna were not included. Our goal here is to understand how data analysis can be made more robust to this class of instrumental response (or any other source of foreground covariance), or alternately how tightly certain instrumental tolerances must be constrained. In Section~\ref{sec:globsig} we review the basic properties of the global signal and describe our foreground model. Section~\ref{sec:simplifiedmodel} builds up the estimator for joint foreground and signal estimation using a simplified model of spectra along independent sight lines. Section~\ref{sec:instrum} considers implications of this model for passband calibration. Section~\ref{sec:spatial_information} develops spatial weights, and Section~\ref{sec:separable_cov} combines the estimators with spatial and spectral weights. Section~\ref{sec:analysis_considerations} describes a number of considerations for using the methods developed here and for global $21$\,cm signal estimation in general. We discuss telescope beam width, the foreground monopole, how aggressively foreground modes should be removed, and susceptibility to Faraday rotation. We also consider mode removal of the pre-reionization absorption feature and extensions of the simple template amplitude constraint considered throughout. We summarize our conclusions in Section~\ref{sec:discussion}. | \label{sec:discussion} Measurements of the global $21$\,cm signal are very challenging due to (1) astrophysical foregrounds and their interaction with instrumental systematics, (2) terrestrial radio interference, and (3) the ionosphere. Here we have examined the first issue, especially in regard to instrumental response. Following $z\sim 1$ $21$\,cm literature \citep{2013MNRAS.434L..46S, 2013ApJ...763L..20M, 2010Natur.466..463C}, we develop a new method where foregrounds can be jointly estimated with the monopole spectrum. This relies on the fact that foregrounds vary across the sky while the cosmological signal is constant. This idea also extends L13, who argue that surveys with moderate angular resolution covering much of the sky are able to better discriminate between the monopole $21$\,cm signal and foregrounds. The key observation arising from our cleaning method is that the instrument should be designed to minimize the generation of new spectral degrees of freedom from the foregrounds. For example, if each line of sight has a slightly different passband calibration, it also requires a new spectral degree of freedom to describe the foregrounds there. In this sense, the instrument ``spreads out'' the variance over modes that ultimately require more aggressive cleaning and signal loss. In contrast, a constant passband calibration error does not increase the rank of the foreground spectral covariance, and its effect is primarily aesthetic. This is fortuitous because obtaining a smooth spectral calibration of an instrument at these frequencies would require very large, expensive structures that are black in radio wavelengths. Simply requiring that galactic foregrounds be smooth to a few percent provides a sufficient passband calibration for signal recovery. In contrast, significant effort must be put into maintaining passband stability to at least $\sim 10^{-4}$. This may require using much of the instrument's sensitivity to integrate against a stable reference. Polarization to intensity leakage is another example of an instrumental systematic that increases the rank of the foreground covariance. It allows spectral oscillations in the Faraday-rotating polarization signal to contaminate the spectral intensity measurement. Because of depolarization, sky polarization is only a fraction of the intensity, and our estimates of the instrumental constraints are more lax, $< 10^{-2}$. In attempting to measure the global $21\,\textrm{cm}$ signal, one is faced with bright contaminating foregrounds that can interact with instrumental systematics in nontrivial ways, overwhelming the cosmological signal that one seeks to measure. In this paper, we have developed methods that bear similarities to various previously proposed intuitive data-analysis techniques. However, our methods arise from a rigorous, self-consistent framework that allows unknown and unanticipated foreground properties to be derived from real data. Such a framework also provides guidance for instrument design. If design requirements can be adequately met, high-significance measurements of the global $21\,\textrm{cm}$ signal will be possible, providing direct access to the rich and complex astrophysics of the first luminous objects and reionization. | 14 | 4 | 1404.7561 |
1404 | 1404.0946_arXiv.txt | We review a system of autonomous differential equations developed in our previous work \cite{DeSantiago:2012nk} describing a flat cosmology filled with a barotropic fluid and a scalar field with a modified kinetic term of the form $\lag=F(X)-V(\phi)$. We analyze the critical points and summarize the conditions to obtain scaling solutions. We consider a set of transformations and show that they leave invariant the equations of motion for the systems in which the scaling solution is obtained, allowing to reduce the number of degrees of freedom. | In cosmology, the scalar fields play an important role due to its broad phenomenology, which can be used in order to describe different phenomena. For that reason they have been used to model phenomena like inflation \cite{Liddle:2000cg}, dark energy \cite{Copeland:2006wr,delaMacorra:1999ff}, dark matter \cite{Magana:2012ph}, bounce scenarios \cite{Peter:2008qz}, and unification models \cite{Bertacca:2010ct}. The canonical Lagrangian for a scalar field is given by $\lag = X- V(\phi)$, where the first term $X=-\frac{1}{2}\partial_\mu \phi \partial^\mu \phi$ is the kinetic term, and the second is the potential that can have different functional forms depending on the model. In the last years, a generalization has been a studied in which the Lagrangian is a general function of $X$ and $\phi$. These Lagrangians were used initially in the study of inflation \cite{ArmendarizPicon:1999rj} and dark energy \cite{Chiba:1999ka}, and in a previous work we used this type of Lagrangian to unify inflation, dark matter, and dark energy \cite{DeSantiago:2011qb}. One particular form of this kind of Lagrangian is the sum-separable, in which \begin{equation} \lag=F(X)-V(\phi), \label{lag} \end{equation} and there is a clear separation between the kinetic term $F(X)$ and the potential term $V(\phi)$. This class of Lagrangians has the advantage of being more easily studied than the more general case while still conserving some of the rich phenomenology that is not present in the canonical case. In a previous work we studied the dynamical system of this class of Lagrangians in a flat Friedmann-Lemaitre-Robertson-Walker (FLRW) cosmology filled with the field and an additional barotropic fluid \cite{DeSantiago:2012nk}. This analysis allowed us to find the critical points of the system and its stability, from which we drew conclusions about the general behaviour and evolution of the system. In the present work we will review this analysis and we will add an extra analysis for the case in which the system has non-trivial critical points. We will show that in that case there is a symmetry that allows the degrees of freedom of the system to reduce. This behaviour is similar to what happens in the canonical Lagrangian with exponential potential. This symmetry allows the existence of scaling solutions even when the Lagrangian is not of the form $\lag=Xg(Xe^{\lambda\phi})$, which in the literature \cite{Piazza:2004df} is considered to be the most general form of Lagrangians with this behaviour. | In this work we made an analysis of the dynamical system associated to a flat FLRW cosmology containing a barotropic fluid and a scalar field with Lagrangian of the form $\lag= F(X)-V(\phi)$. We observed that this system can be described by three variables and its equations of motion, from those we defined $x$, $y$ and $\sigma$ and obtained their evolution equations. For Lagrangians with constant values for $\gamma_k$ and $\Gamma$, the three autonomous differential equations (\ref{dxdN} - \ref{DSDN}) are enough to describe the system. For more general Lagrangians however, it is necessary to obtain dynamical equations for the variables $\gamma_k$, $\Gamma$, and $\Xi$, that in turn will include new variables in terms of higher derivatives of the Lagrangian, that will require extra equations of motion. In order to have only the minimal three equation system, in our work \cite{DeSantiago:2012nk} we used power-law kinetic terms that produce $\gamma_k$ constant, and power-law or exponential potentials that correspond to a constant value of $\Gamma$. We obtained the critical points of the system, and studied their existence and stability conditions, which are summarized in table \ref{tab:1}. The critical points (a), (b), (c), (d) can be reduced for the canonical Lagrangian to those studied in previous works \cite{Wands:2008tv, Gumjudpai:2005ry}, meanwhile ($\alpha$), ($\beta$) and ($\gamma$) are new. In our previous work \cite{DeSantiago:2012nk} we showed that for the existence of the critical points (c) and (d), the Lagrangian needs to be either canonical with exponential potential or the sum of two power-law terms that satisfy the relation (\ref{sym}). In the present work we showed that when this relation holds, the equations of motion (\ref{fried1_aux} - \ref{cont_aux}) are invariant under the set of transformations defined by (\ref{transf}). This symmetry allows to reduce the number of degrees of freedom of the system to two. We showed that $x$ and $y$ are invariant under the same transformation (\ref{transf}) and hence are suitable to describe the system. Therefore, it was possible to obtain $\sigma$ as function of $x$ and $y$ in equation (\ref{bla}). Using this equation we found the two dimensional autonomous system (\ref{dxdN2}, \ref{dydN2}). We showed that for the two dimensional system the critical points are the same as the ones obtained in table \ref{tab:1}, but with $\sigma$ as a function of $x$ and $y$. The critical point (c) can be obtained then as a function of the auxiliary variables $\gamma_k$ and $\gamma_m$ only, eliminating the dependence on $\sigma$, however it is not possible to do it analytically. For the critical point (d) the analytical expression can be obtained as we have done in (\ref{xdclosed}). The symmetry transformations of (\ref{transf}) are similar to those of (\ref{transcan}), corresponding to canonical exponential Lagrangians and studied in \cite{Holden:1999hm}. Both transformations allow the systems to be treated with a two dimensional set of dynamical equations. For the canonical exponential Lagrangian the equations get reduced to (\ref{dxdN}, \ref{dydN}) with a constant value of $\sigma$, while for the Lagrangian whose terms satisfy (\ref{sym}), they get reduced to equations (\ref{dxdN2}, \ref{dydN2}). \ack JDS is supported by ININ Grant. \providecommand{\newblock}{} | 14 | 4 | 1404.0946 |
1404 | 1404.0388_arXiv.txt | In this work we investigate if a small fraction of quarks and gluons, which escaped hadronization and survived as a uniformly spread perfect fluid, can play the role of both dark matter and dark energy. This fluid, as developed in \citep{Brilenkov}, is characterized by two main parameters: $\beta$, related to the amount of quarks and gluons which act as dark matter; and $\gamma$, acting as the cosmological constant. We explore the feasibility of this model at cosmological scales using data from type Ia Supernovae (SNeIa), Long Gamma-Ray Bursts (LGRB) and direct observational Hubble data. We find that: (i) in general, $\beta$ cannot be constrained by SNeIa data nor by LGRB or H(z) data; (ii) $\gamma$ can be constrained quite well by all three data sets, contributing with $\approx78\%$ to the energy-matter content; (iii) when a strong prior on (only) baryonic matter is assumed, the two parameters of the model are constrained successfully. | A huge amount of high-quality observational data collected so far has made the acceleration of the universe an indisputable fact \citep{Riess:1998cb,Perlmutter:1998np,Knop:2003iy, Riess:2004nr,Astier:2005qq,Spergel:2003cb,Spergel:2006hy,Tegmark:2003ud,PlanckXVI}. Such unexpected behaviour has been commonly attributed to an unknown entity acting as a counter-gravitating fluid, dark energy (DE), and has motivated the bloom of an impressive amount of cosmological models which may be able to elucidate its nature. So far, the so called $\Lambda$CDM model is the most accepted cosmological model, and it is based on the well known cosmological constant; however, it still suffers from theoretical drawbacks that make it difficult to reach a conclusive consensus. Many theoretical proposals can be found which attempt to throw some light on the cosmic acceleration mystery, either trying to address its very origin, or (more modestly) attempting at a compelling description of the recent history of our accelerated universe. Some proposals are based on scalar fields, either canonical such as quintessence \citep{Wetterich1988,Peebles:1987}, or with weirder features, such as k-essence \citep{Armendariz2000} or phantom \citep{Caldwell2003} models. Others have an extra-dimensional spirit and invoke braneworlds \citep{LisaRandall1999,LisaRandall1999b}. Dark matter (DM) is the other main, yet unknown, component of the Universe, and it is necessary to produce enough gravitational attraction on certain scales crucial to structure formation. Some of the proposals for the description of accelerated cosmologies rely on (phenomenological) unified pictures (so called unified dark matter models) where a unique exotic fluid accounts for the whole dark sector composed by DE and DM. If we specifically refer to unified dark matter models, then let us remind that most of them resort to the generalized Chaplygin gas (GCG) \citep{Kamenshchik2001265,Bilic200217,Bento2002}, but one can find other (also phenomenological) proposals as those in \citep{Bertacca2010,Bertacca2011}. The list of (accelerated) scenarios can be completed with many other cases. But if we use the popularity criterion among those additional proposals, then modifications to the General Relativity Lagrangian stand out \citep{Capozziello:2003tk,Capozziello:2003gx,Carroll:2003wy,Jain:2010ka}. Nevertheless see \citep{Sahni2000,Peebles2003,Padmanabhan2003,Sahni:2005,Copeland,Frieman2008,Li:2011sd,Mortonson:2013zfa} for reviews on DE models, which provide a wide perspective on the topic of current cosmic acceleration in general. In general we have a vast collection of set-ups which are quite different in their underlying physics, and although many of them (including the {\it concordance} $\Lambda$CDM model) have a great compliance with observational data, none of them is full proof. On the other hand, the proposal by \citep{Brilenkov}, which can be considered as part of the stream of unified dark matter models, has been suggested to explain the nature of DM and the present cosmic acceleration. Such suggestion arises from the hypothesis that a small part of quarks and gluons did not yield to hadronization, and resisted either as isolated aggregates of quark-gluon \textit{nuggets} (QNs) or as a perfect fluid in the form of a quark-gluon plasma (QGP) (uniformly spread on cosmological scales). There have been several works scrutinizing and supporting this guess \citep{Witten:1984rs,Applegate:1985qt,Farhi:1984qu,Chandra:1999tr}, and the idea followed that the QNs could be a good candidate for DM. In fact, a recent perturbative analysis \citep{Brilenkov2}, reached the conclusion that compatibility with observations was possible (for a mechanical perspective on this topic see \citep{Eingorn:2012jm,Eingorn:2012dg,Eingorn:2013faa}). In contrast, the QGP perfect fluid has not gathered the same interest. A recent work \citep{Rahaman:2012tu} explored the possibility that the QGP fluid acted as DM in galactic halos concluding that the corresponding rotation curves were reasonable. At the cosmological level, a QGP fluid was first considered to mimic DM in \citep{Brilenkov}. % Clearly, the theoretical perspective makes the quark bag an attractive one, as it opens the door to an answer to the nature of the two dark components of the universe without resorting to exotic physics. In previous reference the compliance with observations was carried out in an inverse approach, assuming facts hinted by observations the necessary properties of the quark models were derived. But it is absolutely mandatory to work reversely, that is, to assume the model and then to contrast its theoretical predictions with the observational data. This has to be done in an statistically proper way, beyond quantitative sketches. A thorough study will allow to ascertain whether the model is worth exploring further. This is precisely the goal of this work: to establish the viability of the quark bag proposal, which attempts to explain DE and DM in a unified fashion. In order to do that, we perform a standard statistical analysis by using the following astrophysical probes: type Ia Supernovae (SNeIa), Long Gamma-Ray Bursts (GRBs) and observational Hubble data. In the next section, we describe briefly the main conclusions of each scenario sketched in \citep{Brilenkov}. In Section III, we present the observational data samples used in our analysis and finally, in Section IV, we describe and discuss our findings. | \begin{itemize} \item there is not a striking evidence in favour of the QGP(I) model as a way to describe both the accelerated expansion and to account for the amount expected of DM; % \item and the assumption of an ad-hoc prior on $\Omega_m$ strongly favours the QGP(I) model, but then this is a weak result as it relies on a strong initial hypothesis (prior). \end{itemize} Therefore, all in all, we conclude that the QGP(I) model does not \textit{naturally} explain cosmological dynamics. | 14 | 4 | 1404.0388 |
1404 | 1404.5108_arXiv.txt | RE J1034+396, a narrow-line Seyfert-1 active galactic nucleus (AGN), is the first example of AGNs that exhibited a nearly coherent quasi-periodic oscillation (QPO) for the data collected by XMM-Newton in 2007. The spectral behaviors and timing properties of the QPO have been studied since its discovery. We present an analysis of the QPO in RE J1034+396 based on the Hilbert-Huang transform (HHT). Comparing other time-frequency analysis methods, the Hilbert spectrum reveals the variation of the QPO period in great detail. Furthermore, the empirical mode decomposition provides band-pass filtered data that can be used in the O -- C and correlation analysis. We suggest that it is better to divide the evolution of the QPO in this observation into three epochs according to their different periodicities. In addition to the periodicities, the correlations between the QPO periods and corresponding mean count rates are also different in these three epochs. Further examining the phase lags in these epochs, we found no significant phase lags between the soft and hard X-ray bands, which is also confirmed in the QPO phase-resolved spectral analysis. Finally, we discuss the indications of current models including a spotted accretion disk, diskoseismology, and oscillation of shock, according to the observed time-frequency and spectral behaviors. | Quasi-periodic oscillations (QPOs), which contain useful information about the inner accretion disk, are commonly observed in black hole X-ray binaries. Active galactic nuclei (AGNs), as upscaling of black hole X-ray binaries, are expected to have QPO phenomenon. However, AGNs were never significantly found to exhibit QPOs until the first detection in a narrow-line Seyfert 1 (NLS1) galaxy RE J1034+396 made by XMM-Newton observation \citep{Gierlinski2008}. A QPO with period of 3733 s was found to be coherent with the data segment after $\sim25$ ks of this XMM-Newton observation at a $5.6 \sigma$ confidence level \citep{Gierlinski2008}. From the observation on bulge stellar velocity dispersion, \citet{Bian2010} estimated that the mass of the supermassive black hole (SMBH) is $(1-4) \times 10^6 M_{\odot}$. This estimation indicated that this QPO corresponds to a high-frequency QPO in the black hole X-ray binary \citep{Gierlinski2008, Middleton2009, Bian2010}. RE J1034+396, as well as other NLS1s, are also known for their strong soft X-ray excess and highly variability \citep{Puchnarewicz2001}. The spectral behavior may also relate to the presence of QPO. Using the rms spectrum, \citet{Middleton2009} concluded that the energy spectrum of RE J1034+396 could be decomposed by a low-temperature Comptonization of disk emission and a hard power law tail. The QPO is dominated by the variability of the hard power law tail, while the soft component remains constant. Furthermore, \citet{Middleton2011} folded the light curve of different energy bands with the QPO period and found a significant soft lag, which can be interpreted by the reprocessing. \citet{Maitra2010} divided the X-ray photons into high and low phases and found that the low phase spectrum exhibits an absorption edge at $\sim 0.86$ keV, which corresponds to a warm absorber lying at $\sim 9 r_g$ for a $4\times 10^6 M_{\odot}$ black hole. QPOs in X-ray binaries result in broad peaks in their power spectra. This broadening is probably caused by the dramatic variation in the QPO frequency or a modulation with a relatively stable period but fragmented in time. Fourier analysis is insufficient to study the time variability of frequencies because the frequencies defined in the Fourier spectra are assumed to be constant over the entire observation. Based on the timing analysis technique, adding a moving window is a straightforward solution to investigate the variation of the frequency. Several similar analysis methods were introduced, such as the dynamic power spectrum \citep{Clarkson2003a}, and spectrogram \citep{Oppenheim1989}. An improved time-frequency analysis method, i.e., the Morlet wavelet analysis, was also invented and widely applied on astronomical time series, especially quasi-periodic modulations in various time scales. For example, \citet{Lachowicz2010} studied the 4 Hz low-frequency QPO in XTE J1550--564 using the wavelet analysis and Matching Pursuit algorithm. They concluded that the QPO is composed of multiple independent oscillations with the same frequencies, but the oscillation was present intermittently. For the case of RE J1034+396, \citet{Czerny2010} found a drift in the QPO central frequency based on the wavelet analysis. A possible $\sim24$ ks time scale of a QPO period variation was also marginally detected but could not be firmly concluded as a periodicity because the data time span was limited. In addition, they also found a positive correlation between the QPO frequencies, which were determined by the frequencies at the peak of wavelet spectrum, and the X-ray fluxes. Another possible method to investigate the variation of QPO period is the Hilbert-Huang Transform (HHT) proposed by \citet{Huang1998}. The HHT has been successfully applied on the superorbital modulation of SMC X-1 \citep{Hu2011}, 11-years sunspot variability \citep{Barnhart2011}, and the search for gravitational waves \citep{Camp2007}. In the HHT, the instantaneous frequency, which is different from that in the Fourier analysis, is defined as the time derivative of phase function. Thus, the Hilbert spectrum can provide us detailed information in both the time and frequency domains. However, the time interval between the samplings must be much shorter than the variability time scale. The fast oscillations in X-ray binaries, such as the kilohertz QPOs and burst QPOs, cannot be analyzed by the HHT with current observatories. Fortunately, the X-ray intensity of RE J1034+396 is high enough to provide us with a sufficiently sampled X-ray light curve that can be analyzed using the HHT. In addition, HHT can provide us the phase function of the QPO. Thus, phase-resolved spectral analysis, as well as the phase lags between folded light curves of different energy bands are also applicable. We present our analysis on the evolution of the QPO period for RE J1034+396 as well as the QPO phase-resolved spectral variation. In Section \ref{obs}, we briefly introduce the observation made by XMM-Newton and the data selection criteria. The time-frequency analysis, including the HHT analysis, O -- C result, issues with phase lags, and relationship between the QPO period and X-ray flux, are presented in Section \ref{analysis}. The spectral model as well as the phase-resolved spectral analysis are described in Section \ref{spectral}. We further discuss different possible scenarios, including the spotted accretion disk model, diskoseismology, and shock oscillation model in Section \ref{discussion}. Finally, we summarize our work in Section \ref{summary}. | \label{discussion} The QPO detected in RE J1034+396 is believed to be a high-frequency QPO because of its short periodicity. This type of QPO in X-ray binary systems appears in the very high state, while the accretion disk nearly reaches the innermost region of the central black hole. A variety of mechanisms capable of interpreting the QPO phenomenon were proposed. We discussed three of them according to our discoveries, including the evolution of characteristic periods for each epoch, correlation between the QPO period and mean flux, and variation of spectral components. \subsection{The Spotted Disk Model} The most straightforward model to understand the QPO is that it is caused by the Keplerian motion of a temporary hot spot on the accretion disk \citep{Pechacek2006, Bachetti2010}. In this scenario, the QPO frequency is related to the radius of Keplerian orbit. From the HHT and O -- C analysis, we found that the QPO period changes dramatically even between neighboring cycles. This indicates that the hot spot is not strictly anchored on the accretion disk. Instead, the hot spot is wobbling around a characteristic radius. Furthermore, the evolution of the QPO period can be described by three epochs in this observation. This evolution can be interpreted as the change of characteristic radius caused by the instability of the accretion disk. From the analysis of the relationship between modulation period and the flux, we observed a positive correlation between the QPO period and the flux. This relationship, which is basically agree with that observed in \citet{Czerny2010}, does not agree with the prediction of this model. However, some indications of this model are still worth to be addressed. \citet{Czerny2010} compared the fractional rms amplitude of RE J1034+396 and that of the light curve calculated by a spotted accretion disk \citep{Pechacek2006}. They concluded that the inclination angle should be 2 -- 3 degrees if the hot spot is close to the central black hole. However, not only the flare but also the disk, corona, and probably other materials near the black hole contribute to the X-ray emissions. For example, we obtained three components in the spectral fitting. Those X-ray sources would contribute non-modulated or lower-amplitude X-ray emissions such that the fractional amplitude as well as the inclination angle would be under estimated. \begin{figure}[h] \epsscale{1.2} \plotone{inclination_mass_by_if.eps} \caption{Estimated inclination angles for different black hole masses and QPO periods. The blue and red curves are the cases for QPO period of 3784 and 4083 s, respectively. \label{inclination_mass}} \end{figure} Besides the fractional amplitude, another possible indicator that can be used to estimate the inclination angle is the modulation shape. From the simulated light curve in \citet{Czerny2010} and \citet{Pechacek2006}, the QPO modulation shape became highly non-sinusoidal for higher inclination angles. A non-sinusoidal light curve can be expressed as harmonics in the Fourier analysis; however, we could not observe any significant harmonics in the power spectrum. There are two possible explanations for this. The first is that the inclination angle of RE J1034+396 is low, and the QPO profile is close to sinusoidal such that the harmonics are not significantly detected. Another possibility is that the QPO period varies so dramatically that the harmonics are more strongly depressed than the fundamental. Fortunately, the HHT provides us an indicator to probe the non-sinusoidal level of the data. Unlike harmonics in a Fourier analysis, the non-sinusoidal light curve would produce {\it intra-wave} frequency modulations, i.e., the frequency variation within one cycle \citep{Huang1998} in the instantaneous frequency. The frequency derived from the cycle length in the O -- C analysis can be treated as the {\it inter-wave} frequency modulation, i.e., the frequency variation between cycles. Thus, the intra-wave frequency modulation can be estimated by taking the difference between the instantaneous frequency and inter-wave frequency modulation. We then created simulated light curves according to equations 4 and 10 in \citet{Pechacek2006} for different black hole masses and inclination angles. The orbital radius of the hot spot is calculated using Kepler’s law for orbital periods of 3784 and 4083 s. The amplitude of the intra-wave frequency modulation of a simulated light curve can also be calculated using the Hilbert transform on the simulated light curve. For a mass range $1-4\times10^6$ M$_{\odot}$, we can estimate the inclination angle by matching the intra-wave frequency modulation amplitude of the simulated light curves with the observed one. Figure \ref{inclination_mass} shows the relationship between the estimated inclination angle and black hole mass. We found that the inclination angle was between $\sim 27$ and $\sim 57$ degrees. This estimation of inclination angle is probably an over-estimate because the noise in the light curve would also contribute to the intra-wave modulation. \subsection{Diskoseismology} Relativistic diskoseismology, i.e., the oscillations that are trapped in the inner accretion disk, may also be responsible for the presence of QPOs \citep{Perez1997, Silbergleit2001, Wagoner2001, Ortega2002}. Three oscillation modes, i.e., g-, c-, and p-modes, are investigated. The fundamental frequency of these three modes mainly depend on the mass of the black hole, i.e., $f\propto1/M_{BH}$, where $f$ is the QPO frequency and $M_{BH}$ is the mass of black hole. From \citet{Wagoner2001}, the g-mode is dominated by the gravitational-centrifugal force, in which the oscillation frequency anti-correlates with the flux of accretion disk. The c-mode is caused by the non-radial corrugation on the inner accretion disk, while the oscillation frequency is highly dependent on the spin of black hole. The p-mode is dominated by the restoring-force due to the pressure gradient, in which the oscillation frequency shows a positive correlation with the flux of disk. Only the p-mode shows a positive correlation between the flux and QPO frequency, but it is opposite of our results. Thus, the QPO in RE J1034+396 is not likely caused by the p-mode oscillation. For a black hole with mass of 10 M$_{\odot}$, the g-mode oscillation is between $\sim 70$ to $\sim 110$ Hz for different angular momentums and luminosities \citep[see Figure 1 of][]{Wagoner2001}. If we enlarge the black hole mass to that in RE J1034+396, the g-mode oscillation period is between $\sim900$ and $\sim1400$ s for a $1\times10^6$ M$_{\odot}$ black hole, or between $\sim3600$ and $\sim5700$ s for a $4\times10^6$ M$_{\odot}$ black hole. Thus, the preferred black hole mass is close to the upper limit of the estimation made by \citet{Bian2010} if we assume that the QPO is caused by the g-mode oscillation. However, it is difficult for the slope between the g-mode oscillation period and luminosity to exceed $\sim0.1$ unless the angular momentum of the black hole is extremely high \citep[see equation 3 in][]{Wagoner2001}.Thus, the g-mode oscillation cannot provide the observed slope ($\sim0.3$), although this value is less reliable relative to the correlation coefficient because the observed X-ray flux may also contain the emissions from other components. On the other hand, the c-mode oscillation can produce a larger slope between the flux and QPO frequency in the range of our observed value, but the angular momentum of the black hole should be close to zero \citep[see equation 6 in][]{Wagoner2001}. However, the fundamental c-mode frequency of such a slow-rotating black hole is very low. For example, a 10 M$_{\odot}$ black hole with angular momentum of $a=0.05$ has a fundamental c-mode oscillation frequency of $f\lesssim1$ Hz. This oscillation corresponds to a period longer than $10^5$ s for a $10^6$ M$_{\odot}$ black hole. Thus, if the c-mode oscillation is responsible for the observed QPO, it should be high-order harmonics. We also studied the relationship between the QPO frequency and flux in detail by dividing the evolution of the QPO into three epochs. A change in correlation between epochs was observed: a significant anti-correlation between the frequency and flux was observed in epoch C, but no significant correlation was found in epoch A. This is possibly caused by a change in oscillation mode. A correlation in epoch A was not significantly detected; however, we found that the first three cycles appear to behave similar to the flux drop event in epoch B. If we ignore these three cycle, the remaining cycles in epoch A show a strong anti-correlation between the cycle length and flux with a linear correlation coefficient of $-0.61$. Thus, the QPO in epoch A is probably caused by the p-mode oscillation. After epoch B, i.e., the flux-drop event, the g-mode oscillation dominated the observed QPO phenomenon such that we observed a significant direct correlation between the cycle length and flux. All the discussions above are based on the assumption that the luminosity is less than the Eddington luminosity. This is a necessary condition for the thin-disk assumption. However, the luminosity of RE J1034+396 is probably a super-Eddington \citep{Gierlinski2008, Middleton2009}. Thus, we could not exclude all the possible oscillation modes in a super-Eddington accretion disk. \subsection{Oscillation of Shock} \citet{Das2011} proposed a novel model of QPO to investigate the observed slope between the QPO period and X-ray flux in RE J1034+396. In their model, QPO is caused by the oscillations of the shock formed inside the hot accretion flow. The variation in the QPO period is interpreted as the drifting of shock location. This model successfully interpreted the correlation between the QPO period and X-ray flux in \citet{Czerny2010} with slope of $0.92\pm0.03$. However, we have demonstrated that the slope changes as QPO evolves in this observation. In fact, even the relationship likely changes. For example, the QPO period and flux exhibit no significant correlation in epoch A, or even anti-correlation when we omit the first three cycles that show another possible flux-drop event. Thus, the entire observation cannot be interpreted by this single scenario. | 14 | 4 | 1404.5108 |
1404 | 1404.0800_arXiv.txt | {} { Rotation of meteoroids due to gas drag during the ejection from cometary nucleus has not been studied yet. The aim of this study is to estimate the rotational characteristics of meteoroids after their release from a comet during normal activity. The results can serve as initial conditions for further analyses of subsequent evolution of rotation in the interplanetary space. } { Basic dependence of spin rate on ejection velocity and meteoroid size was determined analytically. A sophisticated numerical model was than applied to meteoroids ejected from 2P/Encke comet. The meteoroid shapes were approximated by polyhedrons with several thousands of surface elements, which have been determined by 3D laser scanning method of 36 Earth rock samples. These samples came from three distinct sets with different origin and shape characteristics such as surface roughness or angularity. Two types of gas-meteoroid interactions and three gas ejection models (leading to very different ejection velocities) were assumed. The rotational characteristics of ejected meteoroid population were obtained by numerical integration of equations of motion with random initial conditions and random shape selection. } { It was proved, that the results do not depend on specific set of shape models and that they are applicable to (unknown) shapes of real meteoroids. A simple relationship between median of meteoroid spin frequencies $\bar{f}$ (Hz), ejection velocities $v_{\rm ej}$ (m\,s$^{-1}$) and sizes $D$ (m) was determined. For diffuse reflection of gas molecules from meteoroid's surface it reads: $\bar{f}\simeq 2\times 10^{-3} v_{\rm ej} D^{-0.88}$, and for specular reflection of gas molecules from meteoroid's surface it is: $\bar{f}\simeq 5\times 10^{-3} v_{\rm ej} D^{-0.88}$. The distribution of spin frequencies is roughly normal in $\log$-scale and it is relatively wide; $2\sigma$-interval can be described as $(0.1, 10)\times \bar{f}$. Most of meteoroids are non-principal axis rotators. The median angle between angular momentum vector and spin vector is $12^\circ$. About $60\%$ of meteoroids rotate in long axis mode. Distribution of angular momentum vectors is not random. They are concentrated in the perpendicular direction with respect to the gas flow direction. These results were determined for 2P/Encke comet, but their validity is general. Attention must be paid if the gravitation of nucleus plays an important role. } {} | From the observations of meteors and bolides, there are several phenomena suggesting that meteoroids rotate. (i) The light-curves of some bright meteors show quasi-periodic brightness variations \citep{Spurnyetal2007}. This phenomenon, which is called {\em flickering}, is sometimes interpreted as a result of rotation of non-symmetric meteoroid \citep[e.g.][]{BeechBrown2000, Beech2001, SpurnyBorovicka2001, BeechIllingworthMurray2003}. The rotational origin of flickering has however been doubted and other explanations, such as autofluctuating mechanism or triboelectric effects, were suggested \citep[e.g.][]{BabadzhanovKonovalova2004, Borovicka2006, SpurnyCeplecha2008, Spurnyetal2012}. (ii) Periodic variations in velocity of Lost City bolide were also interpreted as a result of changing cross-section due to rotation \citep{Ceplecha1996, CeplechaRevelle2005}. (iii) Initial radius of meteor trains \citep{HawkesJones1978} and (iv) non-linear meteor trails \citep{Beech1988} can also be a result of meteoroid rotation, as well as the meteoroid bursting in the atmosphere \citep[e.g.][]{StokanCampbellbrown2014}. Unfortunately, precise and reliable determination of preatmospheric rotation from observations of meteors and bolides represents a significant problem so far. The preatmospheric rotation of meteoroids (and more generally, evolution of rotation in interplanetary space) can be studied theoretically. For such studies, it is necessary to describe the action of the processes that may affect the rotation. It was shown, that the radiative effects more efficiently affect the rotation of meteoroids in space than collisions with dust \citep{OlssonSteel1987}. The asymmetry parameter determined by \citet{Paddack1969} together with the time spent in interplanetary space have been used by many authors for estimates of the spin rate of meteoroids \citep{}, but detailed study which would describe the whole physics of meteoroid rotation in space self-consistently, is still missing. For further modeling of the subsequent spin evolution in interplanetary space, the knowledge of initial rotation, just after the meteoroid birth, is also necessary. For asteroidal meteoroids, which originate as debris from collisions in the Main Belt, the initial rotation can be estimated from results of hypervelocity fragmentation experiments \citep[e.g.][]{Fujiwaraetal1989, Martellietal1994, GiblinFarinella1997, Giblinetal1998}. The majority of shower meteoroids are released from parent cometary nucleus during the normal activity of the comet by gas drag \citep{Whipple1950, Whipple1951}. The gas drag mechanism is connected with sublimation of ice at the surface of the nucleus and acceleration of embedded dust grains and pebbles by gas flow away from the comet. If the meteoroid has irregular shape with some degree of windmill asymmetry, the gas may also accelerate its rotation - similarly as in the simple experiment of \citet{Paddack1969}. Although many authors dealt with the ejection process \citep[e.g.][]{Crifo1995, Jones1995, CrifoRodionov1997, Fulle1997, MaWilliamsChen2002, Molinaetal2008}, the rotation of meteoroids caused by gas drag during the ejection has not been studied yet. The aim of the present study is to fill this gap in our understanding of the meteoroids' rotation and to estimate the rotation characteristics of the meteoroids after the ejection from the parent cometary nucleus. It is the first in the assumed series of articles devoted to rotation of meteoroids. In Sec.~\ref{analSec} a simple analytical theory is described, Sec.~\ref{numSec} is devoted to the description of a sophisticated numerical model. The results from the numerical model can be found in Sec.~\ref{resSec} | The present model is not able to determine specific value of meteoroid spin frequency after ejection from 2P/Encke comet due to the lack of reliable ejection velocity data. The dependence of median spin frequency on ejection velocity and meteoroid size (\ref{resultEq}) is however common for all three gas ejection model, despite of very different ejection velocities. The reliable estimate of the spin frequency therefore depends on reliable value of ejection velocity. Equation (\ref{resultEq}) can be rewritten into more simple form as \eq{\bar{f}\simeq 2\times 10^{-3} v_{\rm ej} D^{-0.88}} for diffuse reflection of gas molecules, and \eq{\bar{f}\simeq 5\times 10^{-3} v_{\rm ej} D^{-0.88}} for specular reflection of gas molecules from meteoroid's surface ($\bar{f}$ in Hz, $v_{\rm ej}$ in m\,s\mj, and $D$ in m). But the direct usage of (\ref{resultEq}) for estimates of preatmospheric spin rate of meteoroids is doubtful. During the time, which meteoroid spent in the interplanetary space, the rotation is affected by several phenomena. In the studied size range, the most important are radiative effects \citep{OlssonSteel1987}, i.e. windmill effect and YORP. The timescale of YORP evolution can be estimated by re-scaling of mean doubling time $t_{\rm d}=14$\,Myr, which was determined by \citet{CapekVokrouhlicky2004} for 2-km gaussian random spheres with spin period of 6\,hours on circular orbit with semimajor axis 2.5\,AU, assuming principal axis rotation in asymptotic states. Corresponding values are $\sim 4$ years for $1$\,mm (50\,Hz) and $70$ years for $10$\,cm (0.1\,Hz) Taurid meteoroids. The actual timescales will be however longer due to (i) heat diffusion through the volume of such small bodies \citep{Breiteretal2010}, (ii) evolution of the spin axis direction by YORP effect and (iii) non-principal axis rotation of the most of meteoroids. In any case, the preatmospheric rotation may correspond to the initial rotational state only in for short time spent in the interplanetary space, but this subject is beyond the scope of this article and it will be studied in the future. | 14 | 4 | 1404.0800 |
1404 | 1404.7427_arXiv.txt | Young stellar object observations suggest that some jets rotate in the opposite direction with respect to their disk. In a recent study, \cite{Sautyetal12} have shown that this does not contradict the magnetocentrifugal mechanism that is believed to launch such outflows. Signatures of motions transverse to the jet axis and in opposite directions have recently been measured in M87 (\citeauthor{Meyeretal13}, \citeyear{Meyeretal13}). One possible interpretation of this motion is the one of counter rotating knots. Here, we extend our previous analytical derivation of counter-rotation to relativistic jets, demonstrating that counter-rotation can indeed take place under rather general conditions. We show that both the magnetic field and a non-negligible enthalpy are necessary at the origin of counter-rotating outflows, and that the effect is associated with a transfer of energy flux from the matter to the electromagnetic field. This can be realized in three cases : if a decreasing enthalpy causes an increase of the Poynting flux, if the flow decelerates, or, if strong gradients of the magnetic field are present. An illustration of the involved mechanism is given by an example of relativistic MHD jet simulation. | In a previous Letter \citep{Sautyetal12} we have established that counter rotation in jets from young stars could be a natural consequence of the MHD equations ruling the plasma. We have shown that deceleration of the jet or shocks can induce counter rotation. We have verified it analytically and numerically. In young stars, it is possible to measure observationally the rotation speed of the jet. Counter-rotation has been observed in some cases but remains under debate. However, in the light of our criterion, it is clear that it does not contradict the magnetorotational launching of the jet. In the context of relativistic outflows such as AGN jets, the criterion, if it can be extended, may induce crucial observational consequences. Those jets are well known to be magnetically launched as well. Recent measurements of the polarization and the VLBI Faraday rotation (e.g. \citealp{Mahmudetal13, Algabaetal13}) have confirmed the helicoidal nature of the jet magnetic field which supports the idea of magnetic launching. Faraday rotation measures provide a direct evidence of the magnetic field structure within the jet \citep{Gabuzda03} and allow to confront observations with simulations \citep{BroderickMcKinney10}. Besides, \cite{Meyeretal13} have measured transverse proper motions of knots in the jet of M87. They infer that knots A and C seem to have opposite velocities transverse to the jet which they interpret as counter rotating shocks. They claim that this is consistent with the model of quad relativistic MHD shocks of \cite{Nakamuraetal10}. This model was used to interpret the helical magnetic structure inferred from polarization measurements for the same knots by \cite{Algabaetal13}. It is worth to note that these observations are rather difficult and need a very long time survey. \cite{KVKB09} have shown numerically that counter-rotation appears in their simulations (hot jet of model B2H) as the result of transfer of angular momentum from the fluid to the magnetic component. \cite{Nakamuraetal10} have also shown that the complex structure of the quad relativistic MHD shock model exhibits a reverse shock that flows upstream, rotating in the direction opposite to the forward shock. We show in this Letter a straightforward extension in the relativistic regime of our criterion for counter rotation in young stars which can also be interpreted in terms of the flow energetics. This criterion applies to shock models as well as simulations where the angular momentum and isorotation frequency are conserved. This does not necessarily means it applies to quad shock models. The compatibility with the Riemann problem solution still needs to be checked. However, the present criterion involved a simpler geometry and does not rely on the presence of kink instabilities or precession of the jet axis. Then, even though rotation measurements in relativistic sources are still out of reach, counter-rotation may have strong observational signatures related to the magnetic field structures. In fact, precise measurements of the magnetic field gradients would be needed to get constraints on the jet dynamics and rotation. | In \cite{Sautyetal12}, as well as in this Letter we have shown that counter-rotation is a signature of the magnetization of a jet, both in the classical and the relativistic regimes. Counter-rotation is possible in the following three cases, the first two ones refer to the geometry of the flow while the third one is related to the energetics of the jet, \begin{enumerate} \item Gradients of the magnetic field are associated with a compression of the flow and a sufficiently small $\delta S$ such that Eq.~(\ref{Vphidem3}) is satisfied. \item In a smooth flow in which $\delta S \propto \varpi^2$, a decrease of $\mu_\mathrm{HYD}$ may lead to the inequality~(\ref{Vphidem3}). This can happen if the flow is decelerated. \item In an ultrarelativistic accelerated or constant flow with $\delta S \propto \varpi^2$, part of the decreasing enthalpy flux could be transfered to the electromagnetic field, see Sect.~\ref{sec:energetics}. Then the hydrodynamic part of the energy to mass flux ratio $\mu_\mathrm{HYD}$ should be smaller than its value at the Alfv\'en surface to obtain counter-rotation. \end{enumerate} In all cases, the role of the magnetic field is critical because it provides the agent to absorb the excess of matter angular momentum. Therefore, counter-rotation is only possible in MHD outflows. As explained at the end of Sect.~\ref{sec:energetics} the thermal content in these flows is also important at least near their origin. The first case may be obtained if there are surface instabilities at the edge of the jet. In this case although the jet expands the flux tubes may squeeze. The changes in the size of the flux tube can induce counter rotation. The second case may occur in FRI jets. As a matter of fact, FRI radio galaxies usually exhibit ultra relativistic jets on the parsec scale while the kilo parsec jet is mildly or not relativistic. It is usually interpreted as a strong deceleration of the moving plasma (see \citealp{Melianietal10} and references therein). This strong deceleration may correspond to the second case where we would observe counter-rotation. This is comparable to the non relativistic case studied in \citep{Sautyetal12}. Conversely, if the kilo parsec scale outflow were to be an outer component slower than the inner relativistic spine jet that we see on the parsec scale, provided that both components are accelerated, then it is likely that the jet rotation sense would be the same on all scales. Measuring the rotation of the flow would thus allow to distinguish between a single component decelerating outflow which changes sense of rotation and a two component accelerating jet which does not changes rotation. The third case corresponds to a decrease of the total enthalpy budget at large distances. In such ultrarelativistic outflows, the transformation of thermal energy flux (enthalpy) to Poynting energy flux would induce counter-rotation and a strong increase in the toroidal magnetic field, i.e. matter energy is converted into magnetic energy. The change of sign of the rotational speed would induce an increase of the Poynting flux and of the magnetic component of the total angular momentum. As a byproduct, the toroidal magnetic field would increase as well, creating strong gradients of the toroidal magnetic field in the poloidal direction. At present, the magnetic field direction is measured via the change of Faraday rotation across the jet \citep{Mahmudetal13}. A complex helicoidal magnetic structure in the knots A and C of M87 jet has already been associated with possible counter rotation as we mentioned in the Introduction (\citeauthor{Algabaetal13}, \citeyear{Algabaetal13}, see also \citeauthor{Meyeretal13}, \citeyear{Meyeretal13} and \citeauthor{Nakamuraetal10}, \citeyear{Nakamuraetal10}). Altogether then, also in relativistic outflows counter-rotation does not contradict at all magnetic launching. Instead, it provides another piece of evidence for the essential role played by the magnetic field in the flow. It is important to measure the sense of rotation along the jet, because we may get, among others, an evidence for the energetic and angular momentum interchanges in the jet between its fluid and magnetic parts. | 14 | 4 | 1404.7427 |
1404 | 1404.5614_arXiv.txt | We have spatially resolved five debris disks (HD~30447, HD~35841, HD~141943, HD~191089, and HD~202917) for the first time in near-infrared scattered light by reanalyzing archival \emph{Hubble Space Telescope} (HST)/NICMOS coronagraphic images obtained between 1999 and 2006. One of these disks (HD~202917) was previously resolved at visible wavelengths using HST/Advanced Camera for Surveys. To obtain these new disk images, we performed advanced point-spread function subtraction based on the Karhunen-Lo\`eve Image Projection (KLIP) algorithm on recently reprocessed NICMOS data with improved detector artifact removal (Legacy Archive PSF Library And Circumstellar Environments Legacy program). Three of the disks (HD~30447, HD~35841, and HD~141943) appear edge-on, while the other two (HD~191089 and HD~202917) appear inclined. The inclined disks have been sculpted into rings; in particular, the disk around HD~202917 exhibits strong asymmetries. All five host stars are young (8--40 Myr), nearby (40--100 pc) F and G stars, and one (HD~141943) is a close analog to the young sun during the epoch of terrestrial planet formation. Our discoveries increase the number of debris disks resolved in scattered light from 19 to 23 (a 21\% increase). Given their youth, proximity, and brightness ($V = 7.2$ to 8.5), these targets are excellent candidates for follow-up investigations of planet formation at visible wavelengths using the \emph{HST}/STIS coronagraph, at near-infrared wavelengths with the Gemini Planet Imager (GPI) and Very Large Telescope (VLT)/SPHERE, and at thermal infrared wavelengths with the \emph{James Webb Space Telescope} NIRCam and MIRI coronagraphs. | Infrared surveys have identified more than a thousand nearby star systems for which infrared excesses beyond $\sim 10~\mu$m reveal circumstellar dust produced from the collisional grinding and destruction of small planetesimals. The amount of dust generally decreases with age because small grains are removed by radiation pressure and Poynting-Robertson drag. Without gas to retard their removal, circumstellar grains typically possess lifetimes of 10,000 yr, which is significantly shorter than the age of the star. These circumstances suggest that the grains are replenished from a reservoir of unseen planetesimals such as asteroids or comets, which are perturbed into orbits that lead to dust-generating collisions. Coronagraphic imaging of debris disks has revealed structures such as localized brightness peaks, asymmetries, and warps that suggest the presence of formed or forming planets \citep{Wyatt:2008p2687}. The first three systems with directly imaged exoplanets (Fomalhaut, HR~8799, and $\beta$~Pictoris; \citealp{2008Sci...322.1345K,Marois:2008p2921,Lagrange:2010p3211}) are all stars previously known to host debris disks. For instance, recent ground-based, high-contrast imaging has discovered a $9 \pm 3~M_{Jup}$ planet orbiting 8--13~AU from $\beta$~Pictoris that is consistent with a warp or secondary disk observed in scattered light \citep{Golimowski:2006p501,Lagrange:2010p3211}. Modeling the dynamical effects inferred from scattered-light morphologies places constraints on the architectures of exoplanetary systems \citep{2006ApJ...648..652S}. Models based solely on spectral-energy distributions (SEDs) are inherently degenerate, so direct images are essential to locate the dust and planetesimal belts (the analogs of our asteroid and Kuiper belts) unambiguously. Furthermore, models based only on assumptions about grain sizes and compositions yield disk radii and dust masses that may vary by an order of magnitude \citep{2006ApJ...638.1070H}. However, multi-band scattered-light images of the disks provide measurements of the colors, phase functions, and albedos of the dust that can be used by the models to constrain the physical properties of the dust grains \citep{Golimowski:2006p501,2007ApJ...654..595G,2009ApJ...696.2126S}. Only 23 debris disks have published spatially resolved detections in scattered light as of early 2014, primarily using coronagraphs aboard the \emph{Hubble Space Telescope (HST)} \citep[e.g.,][]{Weinberger:1999p175,Schneider:1999p532,Clampin:2003p118,Ardila:2004p451,Krist:2005p511,Golimowski:2006p501,SSH06,Krist07,Krist:2010p3152,2011AJ....142...30G}. This subset of known debris disks comprises the 16 disks listed in Table 1 of \citet{2011AJ....142...30G}, HD 202628 \citep{2012AJ....144...45K}, HIP 79977 \citep{2013ApJ...763L..29T}, and the five disks reported in this {\it Letter}. Scattered-light images typically possess higher angular resolution than images in thermal emission, and their increased sensitivity to micron-sized particles at large distances from the star provides more detailed information about the spatial distribution of the smallest dust grains. Together, scattered-light and thermal-emission images can be used to constrain the azimuthal dependence of the dust density distribution and the properties of the constituent grains (e.g., composition, size, porosity). The principal difficulty with obtaining such images of disks or planets -- even with the highly stable \emph{HST} coronagraphs or the latest generation of specialized ground-based instruments (GPI, SPHERE, Project 1640, HiCIAO) \citep{2008SPIE.7015E..31M,2008SPIE.7014E..41B,2008SPIE.7014E..42H,2011PASP..123...74H} or the \emph{James Webb Space Telescope (JWST)} NIRCam and MIRI coronagraphs \citep{2010PASP..122..162B} -- is the large contrast between the faint scattered light from the disk and the much brighter halo of starlight from the instrumental point-spread function (PSF). The residual starlight must be precisely calibrated and subtracted during image processing. Recent advances in coronagraphic image processing have been made through the development of sophisticated algorithms for removing the residual diffracted light, which utilize large libraries of reference coronagraphic stellar PSFs \citep[e.g., ][]{Lafreniere:2007p274,LMD09}). Recently, we initiated the Archival Legacy Investigation of Circumstellar Environments (ALICE) project to reprocess comprehensively and consistently archived images from various \emph{HST}/NICMOS coronagraphic surveys for faint circumstellar companions. This program uses the Karhunen-Lo\`eve Image Projection (KLIP) algorithm \citep{Soummer:2012p3139} and the large number of reference stars available in the \emph{HST} archive. KLIP not only decreases computational costs incurred from previous optimization methods, but also improves the ability to detect circumstellar disks by greatly mitigating the disk self-subtraction problem that plagued earlier algorithms. Our ALICE pipeline improves the coronagraphic detection limit for point sources by at least an order of magnitude over classical PSF subtraction methods. For disk images, most of the improvement is obtained at small separations, as demonstrated in Figure \ref{Fig:KLIP} for the previously imaged debris disk around HD~181327 \citep{SSH06}. In this \emph{Letter}, we report some of the first results from ALICE: spatially resolved images of five debris disks that were previously undetected in archived NICMOS data. Four of these disks have never before been imaged in scattered light. The disk around HD~202917 was previously detected by the \emph{HST} Advanced Camera for Surveys (ACS) Guaranteed Time Observer team and reported by \citet{Krist07} at the 2007 Spirit of Lyot conference. The images and analyses presented here constitute ``first looks" at these systems. Each image warrants more in depth analysis, and detailed follow-up papers are in preparation. \begin{figure}[htbp] \center \resizebox{\hsize}{!}{\includegraphics{Fig1.pdf}} \caption{Improved residual starlight subtraction using the KLIP algorithm in the ALICE pipeline for the well-known debris disk around HD~181327. The left image was produced using conventional subtraction techniques \citep{SSH06}, while the right image is obtained using KLIP \citep{Soummer:2012p3139} and a library of reference PSFs from the LAPLACE Archive \citep{SSS10}. KLIP significantly improves the subtraction within 1\farcs5 of the star, as evidenced by the lower residuals within the HD~181327 ring. The newly detected disks presented in this paper lie within this range of angular separation, which explains why they were not seen previously in the NICMOS images. At an angular separation of 1 arcsec, KLIP improves the coronagraphic image contrast by a factor of $\sim50$ over classical PSF subtraction, based on average results from over 7 different coronagraphic images of stars without known circumstellar disks. The reduced apparent surface brightness with KLIP (algorithm throughput) can be calibrated with forward modeling \citep{Soummer:2012p3139}. }\label{Fig:KLIP} \end{figure} | We have obtained new scattered-light images of five debris disks found in the HST NICMOS coronagraphic archive after reprocessing those data with the KLIP algorithm in our ALICE pipeline. Preliminary descriptions of the disks' characteristics are given based on scattered-light image morphology, stellar properties, and SED modeling. More thorough analyses including numerical modeling of disk physical properties will be reported in future papers. Followup observations of these newly seen young disk systems around roughly solar type stars may help elucidate the dynamical processes at work at ages at which terrestrial planets may be forming. Complementary visible-light imaging of the disks around HD~30447, HD~35841, HD~191089, and HD~141943 is being obtained in \emph{HST} Cycle 21 using the Space Telescope Imaging Spectrograph (STIS) coronagraph (M. Perrin, PI), and we will be obtaining further infrared observations using the Gemini Planet Imager. Just as the presence of the exoplanets around HR~8799 and $\beta$~Pic was initially inferred from their infrared excesses due to dust, we recognize a significant likelihood that massive exoplanets exist around these stars as well. | 14 | 4 | 1404.5614 |
1404 | 1404.2202_arXiv.txt | In this paper, we investigate the external field effect in the context of the MOdified Newtonian Dynamics (MOND) on the surface brightness and velocity dispersion profiles of globular clusters (GCs). Using N-MODY, which is an N-body simulation code with a MOND potential solver, we show that the general effect of the external field for diffuse clusters, which obey MOND in most of their parts, is that it pushes the dynamics towards the Newtonian regime. On the other hand, for more compact clusters, which are essentially Newtonian in their inner parts, the external field is effective mainly in the outer parts of compact clusters. As a case study, we then choose the remote Galactic GC NGC 2419. By varying the cluster mass, half-light radius, and mass-to-light ratio we aim to find a model that will reproduce the observational data most effectively, using N-MODY. We find that even if we take the Galactic external field into account, a Newtonian Plummer sphere represents the observational data better than MOND to an order of magnitude in terms of the total $\chi^2$ of surface brightness and velocity dispersion. \\ | The MOdified Newtonian Dynamics (MOND), proposed by Milgrom in 1983 (Milgrom 1983a) is one of the most serious rivals of the dark matter paradigm. It is a phenomenological theory which starts from a very simple ad hoc scaling assumption on the gravitational acceleration $g$. According to this assumption, for accelerations comparable to or less than a critical acceleration $a_0$, $g$ is greater than the Newtonian acceleration $g_N$ in a subtle way. This mild modification originally was innovated to explain the flat rotation curves of spirals, and it proved to be highly successful to get this target (e.g. \citealp{san98}). The problems of the violation of energy and angular momentum conservation for non-spherical mass configurations in MOND \citep{fel84} were immediately fixed by presenting a Lagrangian formalism for it \citep{bek84}. Also, the covariant form of MOND was provided by Bekenstein who introduced a vector field, a scalar field and the gravitational tensor field, the so-called TeVeS \citep{bek04}, to explain relativistic effects such as gravitational lensing. Besides the rotation curves, this theory has had remarkable success in fitting to some other important observations of the local universe. However, MOND faces serious challenges on extragalactic scales. For example it can not completely explain the velocity of galaxies in clusters of galaxies (\citealp{ger92}; \citealp{san94}; \citealp{san99}; \citealp{agu01}). Also, the notorious offset between the baryonic and lensing masses in some galaxy cluster mergers \citep{clo06} has no convincing explanation by MOND (for a review see e.g., \citealp{fam12}). As a matter of fact, MOND has been devised to substitute the assumption of dark matter. As Milgrom himself correctly recognized, his innovation would be in serious trouble in dynamical systems with no mass discrepancy but with internal accelerations $g\lesssim a_0$, e.g., open clusters \citep{mil83a}. To remedy this problem he designed his prescription in a way so that the Galactic gravitational field could suppress the extra acceleration of MOND (the so-called External Field Effect, EFE). Thus, besides checking whether MOND can properly play the role of dark matter in, say, spiral galaxies, we can verify it as the correct dynamics in pure Newtonian systems. To do this, one should choose objects with low mass-to-light ratios and small internal accelerations comparable to $a_0$. Diffuse Galactic globular clusters (GCs) seem to be the best candidates \citep{bau05}. However, as we will elaborate in the next section, if the Galactic field is much higher than $a_0$ the internal dynamics will be Newtonian, regardless of the internal field. Therefore, distant Galactic GCs in the outer halo are the best candidates. Meanwhile, to study the EFE, which is absent in Newtonian dynamics, but is a trait of MOND (and every nonlinear dynamics), the gravitational field in which the GC is embedded should not be negligible. So a limited range with just a few Galactic GCs is at our disposal until the invention of telescopes and spectrographs with higher resolution powers will allow us to investigate kinematics of GCs in the halos of other neighbor galaxies. Several studies have endeavored to discriminate between Newtonian dynamics and MOND using Galactic GCs (e.g., Jordi et al. 2009; Gentile et al. 2010; Baumgardt et al. 2005; Sollima et al. 2010, 2012; Haghi et al. 2009, 2011; Lane et al. 2009; Scarpa et al. 2011). However, until recently most of them used their overall (average) dynamical properties. With the development of high-resolution spectrographs, the velocity dispersion profiles of some GCs are curently available. Thus, more exquisite details of different models can be checked against the observational data. The GC NGC 2419 has recently received a great deal of attention in this regard since it has all the above-mentioned qualifications \citep{bau09}. It has also triggered a hot debate between the MONDian and Newtonian blocks (Ibata et al. 2011a, 2011b; Sanders 2012a, 2012b; for more details see Sect. 5). However, they all regarded this GC as isolated, i.e., with no external field. It is worth noting that Ibata et al. (2011a, hereafter I11a) deduced the ineffectiveness of the external field in alliviating the mismatch of MOND and the observational data of NGC 2419, using MONDian N-body simulation. Nevertheless, one should note that they performed this by applying the external field on the same best-fit models that they found in isolation. In this paper we first investigate the effect of introducing the external field in the MONDian dynamics on the internal dynamics of GCs as a whole. Then we choose the Galactic GC NGC 2419 to check the EFE by finding the best-fit simulated model, having included the external field. In section 2 a brief exposition of MOND and the external field effect is presented. Section 3 describes the simulation approach that is used to obtain dynamical predictions from the model. In section 4, a typical GC is modeled as a spherical collisionless system of stars with negligible binary fraction, and its internal dynamics is studied for a specific mass but different internal and external gravitational fields. Section 5 introduces NGC 2419 and fits a simulated model to its data. Section 6 presents our conclusions and describes caveats one should be mindful of. | In this paper we studied the velocity dispersion and surface brightness of simulated GCs in both strong and weak external fields in MOND. We showed that the MOND dynamics differs greatly from the Newtonian one in GCs. In general, the velocity dispersion of a typical GC in MOND is larger than that predicted by the Newtonian dynamics, especially in the outer regions. A uniform external field causes the internal kinematics to drift toward the Newtonian regime, overwriting the flattening of the velocity dispersion profile at large radii \footnote{Note that such a flat velocity dispersion profile is expected only for models without tidal truncation like the Plummer models adopted here.}. These effects are more spectacular in diffuse GCs (where $g_{int}\ll a_0$), while in compact GCs they are apparent only in the outer regions. \\ In terms of the overall velocity dispersion, the simulated dynamics is in harmony with MOND and its asymptotic behaviors (see Section 2). For every fixed external field, increasing the internal field causes the velocity dispersion to approach the Newtonian value. On the other hand, the same occurs if we fix the internal field and increase the external field. Nevertheless, to explain the internal dynamics of the Galactic GC NGC 2419, introducing the Galactic external field in the model may not be in MOND's favor. In terms of $\chi^2$ goodness-of-fit, our best Newtonian Plummer model is better than the best model in MOND with an external field to an order of magnitude. So this disproves the effectiveness of EFE in MOND in matching it to the observations. Our analysis showed that the main change occurs in the outer parts of the GC, as might have been conjectured, because the external field can show up itself where the star accelerations are much smaller than $a_0$. Meanwhile, the mass of the best Newtonian model ($9.0\times10^5 M_{\odot}$) and its half-light radius ($25$ pc) are comparable with $9.12\times10^5 M_{\odot}$ (obtained by I11a) and $23$ pc (estimated by Harris 1996), respectively, and its $M_*/L$ ($1.9$) is marginally in accord with the constraints found recently by Bellazzini et al.(2012). Therefore, NGC2419 is similar to the other outer Galactic halo clusters Pal 14 \citep{jor09} and Pal 4 \citep{fra12} in that it is difficult to explain their internal dynamics with MOND. However, we should keep some caveats in mind. The Plummer model may not be the best one to describe NGC 2419. Other suitable models might be checked. The chosen interpolating function $\mu(x)$ may not be the best choice. Other forms of interpolating functions may yield different results. Our assumptions should be considered as well. The most critical being that in deriving the surface brightness we assumed a constant mass-to-light ratio, while in general this is not strictly the case, because of the mass segregation in star clusters (Frank et al. 2012; Jordi et al. 2009). So the plausibility of this assumption is debatable and a radially-varying $M_*/L$ with a suitable profile might conclude in a better fit. However, specifically for NGC 2419, I11a argue that the lack of any significant mass segregation, proposed by Dalessandro et al.(2008) on the basis of observational results on the radial distribution of blue straggler stars, shows that the mass-to-light ratio can be reasonably assumed to be constant with radius. Another implicit assumption in our treatment is that we considered ${\bf g}_{ext}$ at the position of NGC 2419 constant in time, equivalent to the assumption of a circular orbit. As a matter of fact, the orbit of this GC is unknown and in the case of an elliptical orbit the Galactic field changes with time and this affects the internal dynamics. However, I11a estimated that the internal dynamical time of this GC is much less than its orbital period. So the assumption of a constant external field equal to the present value is plausible. The effects of non-sphericity and some rotation should also be evaluated. As a GC, NGC 2419 may include some fraction of binaries. Unless the binary star fraction in NGC 2419 is inappreciable, it can inflate the observed velocity dispersion (Cote et al. 2002). However, I11a showed that although the fraction of binaries in NGC 2419 is estimated to be as large as 20\%, their velocity distribution is so peaked around $v=0$ that their effects on the velocity dispersion of NGC 2419 is negligible. At the largest amount, one can invoke the uncertainties on the binary parameters such as mass ratios, orbital eccentricities, period distributions, and mass functions to evaluate their impact on the velocity dispersion. In regards to the EFE, like many other authors, we assumed that the total MONDian acceleration would equal the vectorial sum of internal and external accelerations. However, abandoning this assumption would require a high-resolution MOND simulation with the ability to embrace both Galaxy and cluster distributions to solve the modified Poisson equation. Such a simulation is not available at this time. The assumption of isotropy in the simulation code which is used in this study (N-MODY) might cast a shadow on the results. Certainly not all elliptical structures can be assumed isotropic. In fact, van Albada showed that the formation of stellar systems through dissipationless gravitational collapse leads to isotropic cores and radially-anisotropic envelopes \citep{alb83}. Nowadays, anisotropy attracts more attentions in galactic astrophysics. As an example, a by-product of I11a is that the best Newtonian Michie model is more likely to describe NGC 2419 than the best Newtonian King model by a factor of $10^{118}$. Some changes have to be made in the model-producing programs to include various models of anisotropy, the next step. | 14 | 4 | 1404.2202 |
1404 | 1404.7498.txt | Understanding the dynamics behind black hole state transitions and the changes they reflect in outbursts has become long-standing problem. The X-ray reflection spectrum describes the interaction between the hard X-ray source (the power-law continuum) and the cool accretion disc it illuminates, and thus permits an indirect view of how the two evolve. We present a systematic analysis of the reflection spectrum throughout three outbursts (500+ observations) of the black hole binary GX 339$-$4, representing the largest study applying a self-consistent treatment of reflection to date. Particular attention is payed to the coincident evolution of the power-law and reflection, which can be used to determine the accretion geometry. The hard state is found to be distinctly reflection weak, however the ratio of reflection to power-law gradually increases as the source luminosity rises. In contrast the reflection is found dominate the power-law throughout most of the soft state, with increasing supremacy as the source decays. We discuss potential dynamics driving this, favouring inner disc truncation and decreasing coronal height for the hard and soft states respectively. Evolution of the ionisation parameter, power-law slope and high-energy cut-off also agree with this interpretation. | The most fundamental classification of black hole X-ray binaries (BHXRBs) indicates two distinct classes: persistent and transient systems. Persistent sources are typically wind-fed by high-mass companions, allowing a consistently high accretion rate and hence modest spectral variability. The transient systems however, which have Roche-lobe filling low-mass companions, exhibit dramatic outbursts spanning as many as 8 orders of magnitude in luminosity, but ultimately spend the majority of their existence in quiescence \citep{Remillard06, Fender12}. Remarkably, almost all transients showcase more or less identical characteristic spectral and temporal evolution in outburst, and understanding the mechanisms driving this has become a key focus in astrophysics. At the onset of an outburst the X-ray spectrum is distinctly hard, peaking in power at around 100\,keV, and is well described by power-law ($\Gamma\sim1.6$), hence being commonly known as the low/hard (hereafter `hard') state. In bright hard states an additional soft component is also often observed \citep{DiSalvo01,Miller06,Reis10,Kolehmainen13}, usually attributed to thermal emission with a peak temperature of $\sim0.2$\,keV, whilst there is a high level of aperiodic variability (up to 50\,\%; \citealt{VDK06}). The dominant power-law probably arises from inverse-Compton scattering of `seed' photons, supplied by the accretion disc, in a hot and optically thin `corona' of electrons. Steady radio emission is also observed at GHz frequencies, associated with a steady compact jet \citep{Fender04} and is well correlated with the X-ray emission \citep{Corbel03,Gallo03,Corbel13}. The source spans many decades in X-ray luminosity as it rises up the hard state, but nevertheless shows little spectral evolution. Above a few \% of L$_{\rm edd}$ systems often (but not always) commence a transition to a softer X-ray state. First the source advances into the hard-intermediate state (HIMS), where now the power-law is steeper ($\Gamma\textgreater2$) and the thermal component has increased in both temperature and contribution (now $\sim50\,\%$), resulting in a softer spectrum (see e.g. \citealt{Hiemstra11}). Progressing through this state the flux continually rises whilst the integrated variability begins to decrease. Beyond the HIMS lies the soft-intermediate state (SIMS) which harbours a similar, albeit slightly softer, spectrum to the HIMS. The timing properties are however distinctly different with a fractional RMS now below 10\,\% \citep{Munoz11}. The flux has also continued to increase, and the SIMS typically marks the most luminous phase of the outburst. As the state transition occurs this also reveals contrasting radio emission, by which the jet declines but also displays dramatic flare events often two or more orders of magnitude more luminous than the steady hard state emission (see e.g. \citealt{M-J12,Brocksopp13}). After the brief state transition (typically $\sim10$ days) the system enters the high/soft (hereafter `soft') state, dominated by thermal emission from the accretion disc, which now has an effective temperature of $\sim1$\,keV \citep{Dunn11}. The variability lessens further (now $\textless5\,\%$) whilst the Comptonised emission remains steep and relatively weak throughout. Radio emission is now undetected, and is believed to signify quenching of the jet \citep{Russell11}. The soft state typically lasts for many months, during which there is some variation in hardness, and in some cases one or two brief excursions back into the SIMS, and even sometimes the HIMS. Throughout though the source is generally fading in flux, and upon reaching a few \% in Eddington luminosity makes a state transition through the intermediate states and back into the hard state \citep{Maccarone03} where it fades further into quiescence. The interplay between the respective soft thermal and hard Comptonised emission ultimately defines the two distinct hard and soft states. However, defining the morphology leading to state transitions and separating the two states has proven to be difficult, not least due to an insufficient physical understanding of the corona. Interpreting the role of the jet has also proven to be difficult, even though the X-ray and radio emission correlate well \citep{Corbel03,Gallo03,Corbel13}. Furthermore the base of the jet has been proposed as a source of hard X-rays (e.g. \citealt{Beloborodov99,Markoff05}), heightening the need to understand the connection between the two. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{The reflection spectrum} While photons up-scattered in the corona are observed directly as a hard power-law many will also irradiate the disc leading to a number of reprocessing features, collectively known as the \emph{reflection spectrum} (see \citealt{Fabian10} for a recent review). Fluorescent emission, as atoms de-excite after photoelectric absorption, is of highest prominence and interest in the X-ray band, and is dominated by Fe emission as a result of high abundance and fluorescent yield (which varies as the atomic number $Z^4$). Early works by \cite{George91} and \cite{Matt91} performed Monte Carlo calculations of fluorescent emission resulting in estimates of equivalent widths, line strengths and angular dependance for the Fe K line. Of additional importance is electron scattering which dominates above $\sim$10\,keV (photoelectric absorption dominates below this) and is observed as a peak in flux around 20--40\,keV known as the \emph{Compton hump}. The surface layers of the accretion disc are likely to become ionised by the powerful irradiation arising from the corona, and has lead to many works studying the affect of ionisation upon the reflection spectrum \citep{Ross93,Matt93,Zycki94,Ross99,Nayakshin00,Nayakshin01,Ballantyne01,Ross05}. In particular, \cite{Ross05} represents the grid \textsc{reflionx} which is today the most widely applied reflection model, taking into account the strongest emission lines and self-consistent treatment of the continuum. More recently \cite{Garcia10} introduced \textsc{xillver} (see also \citealt{Garcia11,Garcia13}) which represents a furthering in the treatment of atomic processes, in particular making use of the photo-ionisation code \textsc{xstar}, and is the model applied in this study. The extent of ionisation in the surface layers of the disc is defined by the ionisation parameter $\xi$, which represents the ratio of the illuminating X-ray flux with the gas density \citep{Tarter69,Garcia13}. In BHXRBs lighter elements in the surface layers are expected to be fully stripped by the ionising power of the disc, resulting in a high albedo and values of $\xi$ \citep{Ross93,Zycki94}. This effect also leads to effective Fe emission, since the fully ionised lighter elements are not able to absorb the emitted Fe photons \citep{Matt93}. As it is inherently dependant upon the geometry of the corona producing the photons and the disc intercepting them, the reflection spectrum presents the opportunity to infer changes in the two components in comparison to their own specific emission. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{GX 339$-$4 and this study} GX 339$-$4 is a BHXRB \citep{Hynes03,Munoz08} and one of the most active transient systems, exhibiting numerous outbursts since its discovery \citep{Markert73}, including four complete cycles (with state-changes) in the past twelve years. As a result, GX 339$-$4 is one of the most studied transient systems and over the lifetime of \emph{RXTE}\footnote{Rossi X-ray Timing Explorer} (1995--2012) an extensive archive of data has been amassed, allowing an unparalleled timeline to investigate source variability with the same mission. To this end it has formed the basis of many important works key to our understanding of BHXRBs (see e.g. \citealt{Belloni05,Dunn10,Corbel13} and references therein). Restricting to periods where both the Proportional Counter Array (PCA; \citealt{Jahoda06}) and High Energy X-ray Timing Experiment (HEXTE; \citealt{Rothschild98}) instruments were active this presents three fully sampled outbursts to analyse. In addition GX 339$-$4 is the best monitored BHXRB in the radio band, allowing a unique insight into the outburst nature of transient systems. In this study we examine how the X-ray reflection evolves throughout these three outbursts, presenting one of the most extensive and detailed studies to date of reflection in BHXRBs. We begin by outlining our data reduction strategy in \S , followed by details of the model applied to the observations and our automated fitting procedure in \S\ref{model}. We then present the results of the study in \S\ref{analysis} and then outline and discuss our favoured interpretations in \S\ref{discussion}. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% | \label{discussion} In this work we have utilised three outbursts of GX 339$-$4 to systematically uncover the spectral evolution of black hole binaries. In particular we focus upon the reprocessed reflection spectrum revealing marked changes in the reflection, often not in tandem with the Comptonised emission. As we will discuss now, the most likely explanation for this is distinct evolution of the accretion geometry. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{The overall picture}\label{picture} \begin{figure*} \centering \epsfig{file=figures/BH.eps, width=\textwidth} \caption{An illustration of how geometrical evolution will lead to contrasting changes in the Comptonised and reprocessed emission, and notably an increase in reflection fraction. Hard State (blue): The reflection fraction can be increased by decreasing the inner radius of the accretion disc. For a stable corona this model increases the solid angle subtended by the disc, thus increasing the photons intercepted and reprocessed flux. The Comptonised emission meanwhile remains constant. This interpretation suits the hard state well, where the reflection fraction is low, consistent with the small solid angle of a truncated disc, but gradually increases as the source rises. Soft State: If the scale height of the corona decreases, the portion of Comptonised photons intercepted by the accretion disc will increase, leading to increased reprocessing and thus reflection flux observed at infinity. In addition, the lower height will lead to stronger relativistic effects, ultimately focusing more Comptonised photons onto the disc, and in particular the inner regions. As well as heightened reflection, the Comptonised flux will also decrease as a consequence of the increased anisotropy of the emission. Similar results can be formed with alternative modifications of the corona that ultimately lead to irradiation from a lower mean height. We discuss these models in more details in \S\ref{picture}.} \label{BH} \end{figure*} In Figure \ref{PL_R} we display how the Comptonised and reflection emission co-evolve as the outburst progresses. In the hard state (1) the power-law is typically responsible for 5 times more flux than the reflection ($RF\sim0.2$), however as the source rises the reflection fraction increases yet always remains below $1$. On the other hand, the soft state lies in the main above unity, thus signifying larger levels of reflection than power-law. Furthermore, as the source decays the reflection fraction appears to continue increasing. What makes this interesting is that for a stable geometry the two should vary in tandem as the Comptonised emission feeds the reprocessing observed as the reflection spectrum. Therefore evolution is a signal of changes in the accretion geometry. Figure \ref{BH} illustrates two such transformations that can lead to the observed cycle portrayed in Figure \ref{PL_R}. The truncated disc model has become a popular means to describe the observed characteristics of the hard state, such as the hard spectrum and low reflection fraction (see \citealt{Done07} for a review and references therein). As the inner radius of the accretion decreases the solid angle it subtends beneath the illuminating corona increases. More photons are thus intercepted leading to increased reprocessing and reflected flux. In turn the amount of soft photons entering the corona will increase as the disc moves further towards the black hole leading to cooling of the hot electrons, which is observed through a softer photon index and high energy cut-off (Figures \ref{gamma} and \ref{cutoff}). \cite{Beloborodov99} proposes an alternative explanation for these trends through a `dynamic corona', whereby bulk motion of the emitting plasma reduces the irradiation of the disc. Even with a disc at the ISCO this regime can yield the small reflection fractions observed in the hard state, whilst the beaming will also suppress the soft seed photons, hence retaining a hard spectrum ($\Gamma\sim1.6$). We however observe that both $RF$ and $\Gamma$ increase as the source rises through the hard and into the intermediate states (see \S\ref{PL} and \S\ref{RF}), which would thus require a decreasing bulk motion. This would appear at odds with the favoured internal shocks model outlined in \cite{Fender04}, whereby the jet velocity increases with luminosity/state. Given that there is no deviation from the observed trends in $RF$ and $\Gamma$ would suggest that the illuminating medium remains consistent as well. Furthermore the same process ruling $RF$ and $\Gamma$ is present in the decay of the hard state whereby they both decrease as the source fades ($RF$ fitted slope $1.21\pm{0.06}$), thus in the dynamic corona regime the jet velocity should systematically increase leading to bright radio flares. The reality however is that such flares have only been observed in the rise (\citealt{Fender04}, see also \citealt{Corbel13b}), thus we therefore propose the truncated disc model as our favoured interpretation. High resolution observations with \emph{XMM-Newton} and \emph{Suzaku} have also confirmed that the inner disc radius is truncated and decreases with increasing luminosity in the hard state (\citealt{Plant13,Kolehmainen13}; see also \citealt{Petrucci14}). The lower values for the ionisation parameter recorded in the hard state also agree with a truncated inner disc (see \S\ref{Xi_2}). Furthermore, the reflection component correlates better with the reflection fraction than the Comptonised one (Table \ref{pFlux}), hence this would suggest the changes are driven by the geometry of the disc changing, rather than the corona. The reflection fraction is also almost identical both in magnitude and evolution during the rise and decay (stages 1 and 4) of the hard state (compare Figures \ref{PL_R}b--c) which argues strongly for an identical dynamic in and out of quiescence. In contrast, the reflection dominates the Comptonised emission in almost all of the soft state observations ($RF\,\textgreater1$). Furthermore, as the decay progresses the reflection fraction appears to increase. The scatter in flux due to the reduced hard band signal in the soft state makes fitting the trend difficult, however if the 2002 outburst is excluded a slope of $\alpha=0.49\pm{0.17}$ is resolved (where $\log_{10} S_{\rm PL}=\alpha\log_{10} S_{\rm Ref}+c$). In addition, fitting the SIMS and SIMS-A observations, which track the decay quite well (Figure \ref{HID}), reveals a similar relationship ($\alpha=0.73\pm{0.07}$). The decay of the disc luminosity fits the expected $L_{\rm disc} \propto T_{\rm in}^4$ relation well, hence the inner disc radius is without any reasonable doubt stable in the soft state (see \S\ref{L-T} and also \citealt{Gierlinski04,Steiner10,Dunn11}), thus we can rule out the same mechanism as the hard state. Instead, this acts as a strong indication that the underlying changes modifying the reflection fraction are occurring in the corona. The Comptonsied component correlates better with the reflection fraction than the reflection in the soft and soft-intermediate states, and hence acts as further evidence for this dynamic. In Figure \ref{BH} we illustrate the relationship between the coronal height and the amount of reflection and Comptonised flux observed. A reduction in height increases the portion of Comptonised photons intercepted by the accretion disc, leading to increased reprocessing and thus reflection flux observed at infinity. In addition the lower height will lead to stronger relativistic effects upon each Comptonised photon, ultimately focusing more onto the disc, and in particular the inner regions \citep{Miniutti04,Wilkins12}. This will firstly heighten the reflection flux further due to the increased irradiation. Secondly the Comptonised flux observed will decrease as a result of the increased anisotropy (and tolerance towards the black hole) of the Comptonised emission with lowering height. It is clear to see then how the coronal height can significantly vary the reflection fraction by coupling enhanced reflection with a diminished power-law and we therefore favour this solution as the process behind the increasing reflection fraction in the soft state. While we have showcased the corona as a compact `lamp-post' for illustration purposes, the effect of the coronal height may manifest itself in a different form, such as by varying in extension or collapsing, however the interpretation is still very much alike: the coronal geometry is expected to be evolving as the source decays in the soft state. By using the evolving $RF$ to constrain the accretion geometry we have assumed that the albedo remains constant. This however may not be true since the surface layers of the disc are subjected to a large range of illumination during the different phases of the outburst. The albedo itself is strongly dependant upon the ionisation of the surface layers, hence it must be considered how this may influence our results. Above 10\,keV the scattering cross-section exceeds the absorption cross-section, hence there the albedo of the reflection is close to unity and constant. However, below 10\,keV the albedo is strongly influenced by the ionisation stage: as the layers become more ionised the absorptive opacity decreases, leading to more effective reflection \citep{Ross93, Matt93, Zycki94}. Ultimately a fully ionised layer should act like mirror, corresponding to an albedo of unity. In the case of BHXRBs the surface layers are expected to be kept highly ionised by the hot thermal emission emerging from the disc itself \citep{Ross93}, which is evident by the high ionisation parameter values we record throughout this study (see \S\ref{Xi_1}). Furthermore, the ionisation values we record are large enough that the albedo should remain consistently high, and not far from unity \citep{Zycki94}. We do see some evolution in the ionisation parameter, which particularly contrasting between the hard and soft states (Figure \ref{XI}), but this is not enough to change the albedo by more than a factor of 2 \citep{Zycki94}. As we have described previously, the reflection fraction varies over a much broader range than this in outburst. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{On the nature of the SIMS-A observations}\label{SIMS-A} When classifying spectral states it became apparent that there was two distinct clusters of points with low variability (RMS$\textless5\%$) isolated by their X-ray colour (6--10/3--6\,keV flux; Figure \ref{RMS_hard}). We associate the softer points (X-ray colour $\textless0.175$) with the canonical soft state, whilst the harder points are described by spectra very similar to the SIMS. Additionally, of the 7 observations containing a type-A QPO listed in \cite{Motta11}, 6 are in this group. The seventh observation covered what appears to be part of a transition, hence only a portion of the time-series was used by \cite{Motta11} meaning a RMS of $\sim2\%$ was reported. In our study we utilised the entire observation which raised the RMS to $\sim6\%$. As a result of this association we labelled these points as SIMS-A. The majority of the SIMS-A observations displaying a type-A QPO are close to the top branch of the HID (Figure \ref{HID}), and given that the feature is weak and broad \citep{Motta11} it is likely that many SIMS-A observations where a QPO was not detected simply lack the signal to do so. As stated before the SIMS-A has a very similar spectrum to the SIMS (Table \ref{avpar}). In the HID (Figure \ref{HID}) the SIMS-A appear to be distributed at a slightly higher luminosity than the SIMS, however this distinction becomes more clear in Figure \ref{PL_R}d. The SIMS-A is therefore probably associated with an increase in the Comptonised and reflection components, and is the likely explanation for the hardened X-ray colour. It is also interesting that the SIMS-A remains on the `soft' slope of $\textless1$ (Figure \ref{PL_R}d) describing the general decay of the source, even when there is an apparent increase in power-law and reflection flux each time the source transitions to it. The spectrum of the SIMS-A is dominated by the power-law and reflection, whereas the disc flux is relatively weak (Table \ref{avpar}). In comparison, the disc generally accounts for the majority of flux in the SIMS, as is of course standard for the soft state as well. In Figure \ref{Frac_RMS} we plot the fraction of the total source flux from each component which also emphasises how the disc and Comptonised flux are decreased and increased respectively in the SIMS-A. In total there are 13 instances of transitions to the SIMS-A recorded in our study (i.e. excluding observations already in the SIMS-A). Six of these transitioned from the SIMS, of which 5 lead to an increase in reflection flux and 4 in power-law flux. In 4 of these the disc flux also decreased. Another six observations transitioned from the soft state, all of which are characterised by increases in power-law and reflection flux in addition to a reduced disc flux. Furthermore, in each case the photon index softened and all but one of these transitions lead to a net increase in flux. In the remaining observation the source transitioned directly from the HIMS. It is also interesting that the Comptonised and disc fractions plotted in Figure \ref{Frac_RMS} isolate the SIMS-A observations quite well, whereas the SIMS appears more to be an extension of the soft state. This is typically true for SIMS observations with RMS in the 5--7\,\% range, whilst most of the type-B QPOs reported in \cite{Motta11} correspond to observations with a RMS above this level. We note that the choice of the 5\,\% RMS line separating the SIMS and soft state is rather arbitrary and the true value could be slightly higher. This region is also prone to fast transitions shorter than the typical exposure time of the PCA \citep{Munoz11}. Interestingly, the SIMS-A tend to roughly follow the L$\propto$T$^4$ relation (fitted relation: $L_{\rm disc} \propto T_{\rm in}^{\,4.68\pm{0.18}} (1\sigma)$) commonly attributed to the soft state, but at a lower luminosity (or higher disc temperature), consistent with an spectral hardening factor 1.1--1.3 times larger (Figure \ref{L-T}). An alternative explanation is a reduced disc inner radius, however this is unlikely given that the soft state is well regarded to harbour a disc extending all the way to the ISCO. The inner radius derived from the disc normalisation is also fairly stable and less than that of the soft state. The position on the L--T plot and smaller inner radii are both consistent with the \emph{anomalous} regime of GRO J1655$-$40 \citep{Kubota01} and XTE J1550$-$564 \citep{Kubota04}. In relation to the soft state it is quite easy to explain the harder X-ray colour in the SIMS-A, simply since there is a decrease in the soft (disc) contribution whilst the hard (power-law/reflection) increases. However this fails to explain the consistently low level of variability usually associated with emission from the disc (or lack of Comptonised emission). The SIMS-A do appear to show a similar relationship between the Comptonised fraction and RMS as seen for the SIMS and soft state observations, but offset to a higher fraction. It may be that an additional non-variable component is also present but compensated for by the \textsc{cutoffpl} model. Curiously, in the \emph{anomalous} regime reported by \cite{Kubota01} and \cite{Kubota04} an extra component is required, which is harder than the disc but softer than the power-law, and when accounted for returns the regime to the soft state L--T$^4$ track. This may account for the disparity between the inner radii derived from the reflection and the disc. The same process giving rise to the type-A QPO may be behind the increased hardening of the thermal emission or Comptonised flux as well. We note that the SIMS-A are nevertheless well fitted (Figure \ref{redchi}). \begin{figure} \centering \epsfig{file=figures/FTot_Frac.eps, width=0.42\textwidth} \caption{The fraction of total source flux that comes from the Comptonised (top figure), disc (middle) and reflection (bottom) components respectively. We plot the SIMS, SIMS-A and soft state observations to indicate how the source spectrum is different between these three principle states. Whilst the fractional RMS has proven to be a better indicator of state than spectral hardness, clearly the states are not completely distinct, particularly between the SIMS and soft state (see \S\ref{states}). Note that the plotted reflected fraction is the ratio of the reflection and total source flux, and should not be confused with the $RF$. We calculated each flux in the same band as the RMS (2--15\,keV) to make the two axis directly comparable, however the same trend is seen using 0.1--1000\,keV flux as well.} \label{Frac_RMS} \end{figure} %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{The accuracy of this study and comparisons to other works}\label{accuracy} \begin{figure} \centering \epsfig{file=figures/red_chi.eps, width=0.5\textwidth} \caption{The reduced-$\chi^2$ distribution for each best-fit model in our analysis, with 95\% of observations reporting $\leq1.5$. } \label{redchi} \end{figure} In using \emph{RXTE} our approach is quantitative more than qualitative, hence it is important to examine the accuracy of our study. To do this we can examine how well investigations with high-resolution missions like \emph{XMM-Newton} and \emph{Suzaku} agree with our results. Such studies of the hard state \citep{Reis10,Done10,Plant13,Kolehmainen13} reveal moderate ionisation ($\textless\log{3}$) and a low reflection fraction akin to our results. Studies of the HIMS are sparse due to the short duration of this state, but observations analysed by \cite{Hiemstra11} and \cite{Reis11,Reis12} agree with the evolution outlined in \S\ref{spectra}, in particular an increased level of ionisation ($\textgreater\log{3}$) and reflection fraction (the equivalent width is now 250--450\,keV vs. 50--150\,keV in the hard state). \cite{Kolehmainen11} investigated the SIMS and soft state of GX 339$-$4, confirming consistently high ionisation and increased reflection fraction habitual to the soft state in this study. Furthermore, they discuss difficulties in fitting the Fe line profile in the soft state, which is hampered since the disc is providing most of the flux and additional curvature in the continuum at 6\,keV. This is most likely reason for the large scatter in $r_{\rm in}$ during the soft state (Table \ref{avpar}). We can also compare our results to previous works with \emph{RXTE}. Recently \cite{Reis13} performed a similar analysis with one outburst of the black hole XTE J1650$-$500 finding similar results. For example, they also plot reflection versus power-law flux displaying similar evolution throughout the states (Figure \ref{PL_R}). \cite{Dunn08} also studied reflection in GX 339$-$4, instead fitting the Fe line with a Gaussian. Remarkably, even with this simpler approach, very similar evolution is revealed. In their Figure 8 \cite{Dunn08} plot Fe line flux against 7--20\,keV flux (essentially reflection versus power-law flux) finding the same looped evolution and an increased fraction of reflection to power-law in the softer states. Furthermore, the fitted slopes to their `hard' and `soft' observations are almost identical to those in this study (Figure \ref{PL_R}b--d), whilst the evolution of the equivalent width in their Figure 7 is very similar to the trend of its analogous parameter $RF$ analysed in this investigation (Figure \ref{PL_R}). It is also possible that the three components (disc, power-law and reflection) exclusively assumed are not the full description of the X-ray spectra from BHXRBs. For example, in the AGN community much controversy exists due to the potential effect of absorption, and in particular the level, even in some cases presence, of reflection \citep{MillerL08}. No strong indicator of absorption, such as dips, eclipses or narrow lines has been observed from this source \citep{Ponti12}, although this does not necessarily rule out that absorption is present or influences our results. It should be noted though that the reflection spectrum will also be influenced by photons emitted in the disc \citep{Ross93,Ross07}, and this is not currently accounted for by \textsc{xillver-a} and is most likely to have an effect in the soft state where the disc is very strong. The convolution model \textsc{rfxconv}, which combines the ionised reflection of \textsc{reflionx} \citep{Ross05} and the angle-dependant Compton reflection of \textsc{pexrav} \citep{Magdziarz95}, can accept any input continuum, such as that from a disc. \cite{Kolehmainen11} applied \textsc{rfxconv} to soft and soft-intermediate state \emph{XMM-Newton} observations of GX 339$-$4, which through the soft bandpass could directly infer the disc contribution to the reflection spectrum. The results of \cite{Kolehmainen11} are very similar to ours (see e.g. their Table 3), hence we believe our analysis using \textsc{xillver-a} should be robust to the effects of the disc contribution. However, we do note that self-consistent ionised reflection from illumination by the Comptonised \emph{and} disc emission is needed to confirm our results. Finally the chi-square distribution is excellent, with 95\% of observations having a reduced--$\chi^2$ of $\leq1.5$ (Figure \ref{redchi}). %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% \subsection{Why is the reflection spectrum more ionised in the soft state?}\label{Xi_2} In \S\ref{Xi_1} we analyse the ionisation parameter, defined as the ratio of the illuminating flux and the gas density, where it emerges that the parameter is significant larger in the soft state. This would appear at odds with the state of the illuminating flux, which typically peaks in the brighter stages of the hard state and subsequent transition into the HIMS (Figure \ref{PL_R}). To explain this one important aspect to consider is the increased disc emission in the soft state, which will undoubtedly have an ionising affect on the surface layers in addition to the illumination from the corona above. To this end \cite{Ross07} investigated the impact of the disc, finding that increased thermal radiation and peak temperature will ultimately result in a more ionised spectrum. In particular the Fe profile is strongly affected as a result of higher ionisation stages and greater Compton-broadening. Our chosen reflection model (\textsc{xillver-a}) does not account for change in the thermal emission throughout the outburst and thus is likely to react to changes in the disc by varying the ionisation parameter, in particular with an increase in softer states. Another explanation for the apparent change in illumination could result from varying the area of the disc and thus the solid angle it subtends below the corona. If the inner accretion disc is truncated then the amount of Comptonised photons intercepted by the disc will decrease rapidly. Furthermore, assuming a lamppost geometry, the illumination pattern should roughly go as R$^{-3}$, thus the reprocessed spectrum will be dominated by emission from the most central region, which in turn likely represents the most ionised zone of the disc because of the peaked illumination there. Even with a small level of truncation the inferred ionisation level will probably diminish significantly, and this contrast may also be heightened if light-bending is at play (see Figure \ref{BH}). This argument works well since, as discussed in \S\ref{accuracy}, high-resolution spectroscopy resolves the inner disc to be truncated and of a lower ionisation stage in the hard state (e.g. \citealt{Plant13}). %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% | 14 | 4 | 1404.7498 |
1404 | 1404.4424_arXiv.txt | {The relic density of symmetric and asymmetric dark matter in a Gauss-Bonnet (GB) modified Randall-Sundrum (RS) type II braneworld cosmology is investigated. The existing study of symmetric dark matter in a GB braneworld (Okada and Okada, 2009) found that the expansion rate was reduced compared to that in standard General Relativity (GR), thereby delaying particle freeze-out and resulting in relic abundances which are suppressed by up to $\mathcal{O}(10^{-2})$. This is in direct contrast to the behaviour observed in RS braneworlds where the expansion rate is enhanced and the final relic abundance boosted. However, this finding that relic abundances are suppressed in a GB braneworld is based upon a highly contrived situation in which the GB era evolves directly into a standard GR era, rather than passing through a RS era as is the general situation. This collapse of the RS era requires equating the mass scale $m_{\alpha}$ of the GB modification and the mass scale $m_{\sigma}$ of the brane tension. However, if the GB contribution is to be considered as the lowest order correction from string theory to the RS action, we would expect $m_{\alpha} > m_{\sigma}$. We investigate the effect upon the relic abundance of choosing more realistic values for the ratio $\mathcal{R}_{m} \equiv m_{\alpha}/m_{\sigma}$ and find that the relic abundance can be either enhanced or suppressed by more than two orders of magnitude. However, suppression only occurs for a small range of parameter choices and, overwhelmingly, the predominant situation is that of enhancement as we recover the usual Randall-Sundrum type behaviour in the limit $\mathcal{R}_m \gg 1$. We use the latest observational bound $\Omega_{DM}h^2 = 0.1187 \pm 0.0017$ to constrain the various model parameters and briefly discuss the implications for direct/indirect dark matter detection experiments as well as dark matter particle models.} \begin{document} | \label{sec:intro} Precision astrophysical and cosmological measurements have now established that a significant fraction of the matter content in the universe is composed of non-baryonic Dark Matter (DM)~\cite{Hooper}. The data favour cold (non-relativistic) dark matter (CDM) and give the present density as ($68\%$ C.L.)~\cite{Lahav:2014vza} \begin{equation} \Omega_{DM} = 0.1187\pm 0.0017\,h^{-2},\label{eq:dm_abun} \end{equation} where $\Omega_{DM}$ is the dark matter density as a fraction of the total mass-energy budget and $h = 0.678\pm 0.008$ is defined by the present value of the Hubble constant $H_0 = 100\, h$ km/s/Mpc. The most popular theoretical CDM candidates are WIMPs (Weakly Interacting Massive Particles) with mass $m_{\chi} \sim \mathcal{O}(10 - 1000)$ GeV. One viable WIMP candidate is the neutralino, the lightest supersymmetric particle in supersymmetric extensions of the Standard Model (SM) in which $R$-parity is conserved. The origin of the DM can be explained by the thermal relic scenario~\cite{KandT}: at early times, frequent interactions keep the DM particles in equilibrium with the background cosmic bath. As the universe expands and cools, the Boltzmann suppressed interaction rate drops below the expansion rate and the DM particles fall out of equilibrium. At this point - known as particle freeze-out - both annihilation and creation processes cease and the number density redshifts with expansion; the surviving 'relic' particles constitute the dark matter density we observe today. Due to the Boltzmann suppression factor in the equilibrium number density, the present dark matter abundance depends sensitively on the timing of freeze-out: the longer a species remains in thermal contact with the background bath, the lower its density at freeze-out. In the standard cosmological model of cold DM with a non-zero cosmological constant (denoted the $\Lambda$CDM model), particle freeze-out occurs during the radiation dominated era when the expansion rate $H \sim T^2/M_{\mathrm{Pl}}$ (where $M_{\mathrm{Pl}} = 1.22\times 10^{19}$ GeV is the Planck mass). In this scenario, a DM candidate with a weak scale interaction cross section, $\sigma \sim G_{\mathrm{F}}^2\,m_\chi^2$, freezes out with an abundance that matches the presently observed value~\eqref{eq:dm_abun} - this is known as the 'WIMP miracle' and strongly motivates thermal WIMP dark matter models. Despite the observational success of $\Lambda$CDM, current datasets leave the physics of the universe prior to Big Bang Nucleosynthesis (BBN) ($t \sim 200$ s) relatively unconstrained. If the universe experiences a non-standard expansion law at early times, and in particular during the era of DM decoupling, particle freeze-out may be accelerated (or delayed) and the relic abundance enhanced (or suppressed)~\cite{DBarrow1982501,PhysRevD.81.123522,Pallis:2009ed,Salati2003121,Arbey200846,Gelmini:2013awa,Iminniyaz:2013cla,Meehan:2014zsa} (see also~\cite{Gelmini:2009yh}). An interesting class of alternative cosmological models that address this pre-BBN era is provided by the braneworld scenario in which the observable universe is a 3(+1) dimensional surface (the 'brane') embedded in a five dimensional bulk spacetime. Standard Model particles are confined to the surface of the brane whilst gravity propagates in the higher dimensional bulk~\cite{Langlois:2002bb,Maartens:2010ar}. This class of models is motivated by (super)string theory and M-theory which require additional spacetime dimensions for internal consistency. In the widely studied Randall-Sundrum type II (RSII) model~\cite{Randall:1999vf}, General Relativity (GR) is recovered on the surface of a 3(+1) Minkowski brane located at the ultraviolet boundary of a five dimensional anti-de Sitter bulk. The warped geometry of the bulk spacetime ensures the fifth dimension is only accessible in the ultraviolet regime and that $\Lambda$CDM is reproduced in the low energy limit. Relic DM abundances in a RSII braneworld model have been investigated for both the case of symmetric DM \cite{PhysRevD.70.083531,PhysRevD.71.063535,PhysRevD.73.063518,PhysRevD.79.115023, AbouElDahab:2006wb,Meehan:2014zsa}, in which the DM particles are Majorana fermions, that is the particles $\chi $ and antiparticles $\bar{\chi}$ are identical, $\chi = \bar{\chi}$, and the case of asymmetric DM \cite{Meehan:2014zsa} in which the particles and antiparticles are distinct, $\chi \ne \bar{\chi}$. In both cases the enhanced early time expansion rate boosts the final relic abundance. In this article we consider an extension of the RSII model which incorporates a Gauss-Bonnet (GB) higher order curvature term in the bulk action integral, thus modifying the braneworld dynamics at high energies.\footnote{The inclusion of a GB term affects early universe inflation and modifies both scalar and tensor primordial perturbations and the consistency relation between them~\cite{PhysRevD.70.083525,Tsujikawa2004a,Tsujikawa2004b,Calcagni2013}. Although it produces an enhanced ratio $r$ of the tensor to scalar perturbations~\cite{Neupane:2014}, it is still compatible with the recent Planck~\cite{Ade:2013zuv} and BICEP2~\cite{BICEP2} measurements for the case of single scalar field $m^{2}\phi^{2}$ inflation. For a similar study in the regular Randall-Sundrum model see~\cite{Okada:2014nia}.} The relic density of DM in the Gauss-Bonnet braneworld scenario has been studied by~\cite{PhysRevD.79.103528} for the case of symmetric DM. The GB braneworld effect is treated approximately through the use of a simple multiplicatively modified Hubble expansion which can be interpreted as a multiplicatively modified annihilation cross section in the Boltzmann rate equation and allows the development of an approximate analytic expression for the asymptotic relic abundance. They found that the expansion rate was reduced in the GB model, delaying particle freeze-out and leading to a suppressed relic abundance. This is in direct contrast to the behaviour observed in the RSII braneworld model. This finding, however, is based upon a highly contrived situation in which the Gauss-Bonnet expansion era evolves directly into a standard General Relativity expansion era, rather than passing through a Randall-Sundrum expansion era as is the general case. This collapse of the RS era requires equating the mass scale $m_{\alpha}$ of the GB modification and the mass scale $m_{\sigma}$ of the brane tension. However, if the GB contribution is to be considered as the lowest order correction from string theory to the RS action, we would expect $m_{\alpha} > m_{\sigma}$. It is therefore important to investigate the effect upon the relic abundance of choosing more realistic values for the ratio $\mathcal{R}_{m} \equiv m_{\alpha}/m_{\sigma}$ of these two mass scales. In the present paper we revisit the calculation of the relic abundance of DM in the GB scenario and study the effects of breaking the assumption $\mathcal{R}_{m}=1$ made by \cite{PhysRevD.79.103528}, replacing it by more realistic values. We also extend the investigation to consider both symmetric and asymmetric DM species and discuss the implications for DM detection experiments and DM particle models. In the next section we introduce the action integral for the braneworld bulk which includes the Gauss-Bonnet higher curvature term and discuss the modified Friedmann equation in this model. Then, in section~\ref{sec:sdm}, we calculate the DM relic abundance in the Gauss-Bonnet braneworld scenario before deriving constraints on the GB model parameters using the observed relic density. This is repeated for the case of asymmetric DM in section~\ref{sec:adm} and, finally, in section~\ref{sec:con} we summarize our results. | \label{sec:con} Relic abundance calculations provide an important test of non-standard cosmological scenarios in the early pre-BBN universe (see~\cite{Gelmini:2009yh} for further discussion). In this article we have revisited the relic abundance investigation in the Gauss-Bonnet braneworld scenario in which a Gauss-Bonnet curvature invariant is added to the Randall-Sundrum braneworld action. A previous investigation by~\cite{PhysRevD.79.103528} found that the dark matter density is suppressed in the GB braneworld model, however, this conclusion is based on a highly contrived assumption that collapses the Randall-Sundrum expansion era, leading to a slower early time expansion law. We find that when this assumption is relaxed, the early time expansion rate can be either faster or slower than the standard expansion law, depending on the model parameters. In turn, the dark matter relic abundance is either enhanced or suppressed by up to several orders of magnitude with respect to the standard cosmology result, respectively. Importantly, when realistic parameter values are chosen, the early time expansion rate is typically faster than the standard expansion law during the era of dark matter decoupling and the resulting relic abundance is enhanced. Moreover, in the limit $\beta\lll 1$ (corresponding to $\mathcal{R}_m \gg 1$) the usual Randall-Sundrum type behaviour is recovered~\cite{PhysRevD.70.083531,PhysRevD.79.115023}. We have also investigated the GB braneworld effect on asymmetric dark matter species and found that the enhanced annihilation cross section required to provide the observed relic density is capable of producing an amplified annihilation signal with respect to the symmetric signal in the standard cosmological scenario. This effect, which is contrary to the usual expectation, has also been demonstrated in quintessence, scalar-tensor~\cite{Gelmini:2013awa} and Randall-Sundrum braneworld models~\cite{Meehan:2014zsa}. The implications of the latest Fermi-LAT constraints on the dark matter annihilation cross section have been considered for both the symmetric and asymmetric models. For small $\beta$, corresponding to realistic values for the mass ratio $\mathcal{R}_m$, larger values of $\mu^{2}$ are favoured, suggesting that the Gauss-Bonnet braneworld expansion rate has reduced to the standard expansion law before dark matter decoupling. The present investigation is timely because the weak scale cross section relevant to generic relic abundance calculations should be accessible to the next generation of direct and indirect detection experiments~\cite{Bauer:2013ihz}. Therefore, additional constraints and/or an unexpected signal from these experiments could point to new physics in the era prior to BBN. Our investigation also has implications for dark matter particle models and scans of supersymmetric parameter space. If the early time expansion rate is in fact slower than the standard scenario, particles which are typically overproduced in the standard cosmology and thus ruled out by relic density constraints, may be rescued in the GB scenario. | 14 | 4 | 1404.4424 |
1404 | 1404.1037_arXiv.txt | We present the first detailed study of the RR Lyrae variable population in the Local Group dSph/dIrr transition galaxy, Phoenix, using previously obtained HST/WFPC2 observations of the galaxy. We utilize template light curve fitting routines to obtain best fit light curves for RR Lyrae variables in Phoenix. Our technique has identified 78 highly probable RR Lyrae stars (54 ab-type; 24 c-type) with about 40 additional candidates. We find mean periods for the two populations of $\langle P_{ab}\rangle = 0.60 \pm 0.03$ days and $\langle P_{c}\rangle = 0.353 \pm 0.002$ days. We use the properties of these light curves to extract, among other things, a metallicity distribution function for ab-type RR Lyrae. Our analysis yields a mean metallicity of $\langle [Fe/H]\rangle = -1.68 \pm 0.06$ dex for the RRab stars. From the mean period and metallicity calculated from the ab-type RR Lyrae, we conclude that Phoenix is more likely of intermediate Oosterhoff type; however the morphology of the Bailey diagram for Phoenix RR Lyraes appears similar to that of an Oosterhoff type I system. Using the RRab stars, we also study the chemical enrichment law for Phoenix. We find that our metallicity distribution is reasonably well fitted by a closed-box model. The parameters of this model are compatible with the findings of \citet{hid09} further supporting the idea that Phoenix appears to have been chemically enriched as a closed-box-like system during the early stage of its formation and evolution. | \label{sec:intro} With a distance modulus of $(m - M)_0 = 23.09 \pm 0.10$ mag corresponding to a distance of $d = 415 \pm 19$ kpc and a distance from M31 of $600$ kpc \citep{hid09}, Phoenix presents an opportunity to study the evolution of a dwarf galaxy without significant perturbations exerted by massive galaxies, while still being close enough to obtain good sampling of its stellar population. Since its discovery \citep{sch76}, Phoenix has gone from being classified as a distant globular cluster to its currently accepted state as a dwarf transition (dSph/dIrr) type galaxy, which are characterized by recent star formation but lacking any prominent H II regions \citep{mat98}. This is supported by observations of an H I region near the galaxy, likely due to gas expelled from supernova winds, that appears to be associated with recent ($\le$ 100 Myr) star formation \citep{yng07}. Previously, \citet{mart99} discovered two perpendicular, elliptical components in the structure of Phoenix. The inner ellipse is oriented in the east-west direction and contains the young stars in the galaxy. The outer ellipse is rotated $\sim 90^{\circ}$ from the inner and contains no young stars. This indicates that either star formation has recently occurred exclusively in the center of this dwarf galaxy, or stars have formed in an envelope that shrinks over the time due to the natural reduction of the pressure due to the gas. This last hypothesis was recently suggested in a detailed study of the star formation history (SFH) of Phoenix performed by \citet{hid09}. Its distance from massive galaxies, transition type, associated H I region, and the two perpendicular, elliptical components with distinctly different stellar populations make the SFH of Phoenix particularly interesting. In their work, \citet{hid09} compared synthetic color-magnitude diagrams (CMDs) to the observed CMD of Phoenix in order to derive the star formation rate (SFR) as a function of both time and metallicity. The SFH for the entirety of Phoenix was not fit well by any one standard chemical evolution model (e.g. closed-box, infall, or outflow; see \citet{pag09, pei94} for model details) indicating a relatively complex star formation history over nearly a Hubble time ($\sim 13$ Gyr). However, they suggest that a closed-box model is compatible with the SFH of Phoenix until about 6-7 Gyr ago when it appears to have experienced a sudden burst of chemical enrichment. Thus, an independent measurement of the abundances of stars that formed during this early epoch probing the galaxy's chemical evolution at that time could test the validity of this analysis. In this work, we study this early chemical evolution using the RR Lyrae stars present in Phoenix. The RR Lyrae stars are pulsating horizontal branch (HB) stars in the instability strip. They are observed to pulsate in three modes. The ab-types (RRab) pulsate in the fundamental mode; the c-types (RRc) pulsate in the first overtone, while the d-types present both the fundamental and first overtone modes of pulsation. The discovery of RR Lyrae stars in a system indicates the presence of an old stellar population ($\gtrsim 10$ Gyr, \citet{smi95}) characteristic of their low masses ($\approx 0.6 M_{\odot}$). Thus, in analyzing their properties one can probe the conditions of the system at these early epochs. Extensive studies of RR Lyrae stars have uncovered many relations between their pulsation properties and useful astrophysical quantities \citep{san93, fer98, san06, jur96, alc00, mor07, nem13, gul05}. Among these, there is a relation between periods, amplitudes, and metallicities of RRab stars \citep{alc00}. In particular, they define a reduced period, $log$ PA $= log P + 0.15 A_V$, and found that the iron abundances, [Fe/H], of the Galactic globular clusters M3, M5, and M15 correlate with this reduced period. This provides a straightforward method for deriving the metallicity distribution function (MDF) for the RR Lyrae population in a system. \citet{gal04} previously investigated the variable star population within Phoenix. Specifically, they observed the coexistence of anomalous and short-period classical Cepheid variables, as well as identified a previously undetected population of RR Lyrae candidates within the galaxy. That study is the first and only detection of RR Lyrae stars in Phoenix, but due to observational constraints (relatively high photometric errors compared with RR Lyrae pulsation amplitudes), it could not provide an analysis of this population. In this work, we present the first in-depth study of the RR Lyrae population in Phoenix, increasing the number of highly probable RR Lyrae stars with light curve properties by a factor of $\sim 20$. We analyze the properties of this RR Lyrae population with the goal of shedding light on the early evolutionary history of Phoenix. This paper is organized in the following manner. Section \ref{sec:obsred} discusses the observations used in this study and how these data were reduced. Section \ref{sec:var} describes how variable star candidates were selected and characterized as well as how the artificial RR Lyrae simulations were performed in order to characterize biases inherent in our analysis. In Section \ref{sec:comp}, we compare our RR Lyrae sample with candidates identified in \citet{gal04}. In Sections \ref{sec:res} and \ref{sec:disc}, our results are presented and discussed. Finally, Section \ref{sec:conc} summarizes the conclusions drawn from this work. | \label{sec:conc} We have presented the first detailed study of the RR Lyrae stars populating the Phoenix dwarf galaxy. We used light curve template fitting routines to identify and characterize RR Lyrae variables within Phoenix using archival WFPC2 data. The cadence and phase coverage of these data coupled with our fitting routines allowed us to increase the number of highly probable RR Lyrae stars to $78$, with $\sim 40$ more RR Lyrae candidates observed outside of the instability strip. The mean periods calculated for the two types of RR Lyrae found in Phoenix are $\langle P_{ab}\rangle = 0.595 \pm 0.032$ d and $\langle P_{ab}\rangle = 0.353 \pm 0.002$ d for the ab- and c-types, respectively. We have used the properties of the RR Lyrae population within Phoenix to probe its behavior at early times. Using the best fit light curve properties, we have constructed the Bailey diagram for the RR Lyrae stars in Phoenix which displays the RRab stars apparently following the OoI relation. However, the position of Phoenix in the [Fe/H]-$\langle P_{ab}\rangle$ plane lies within the Oosterhoff gap, indicating that it is likely of intermediate Oosterhoff type. This is consistent with most other Milky Way dwarf satellite galaxies, however the discrepancy between the Bailey diagram and the calculated values of $[Fe/H]$ and $\langle P_{ab}\rangle$ warrants further investigation. We used the minimum light colors of the RRab stars in Phoenix to estimate the line of sight reddening to the galaxy. Using this method, we calculated the reddening to Phoenix to be $E(V-I) = 0.07 \pm 0.06$ mag. This does not agree well with the \citet{sch98} maps. However, this does qualitatively agree with previous determinations of the reddening to Phoenix indicating internal sources of extinction resulting in such a discrepancy. We also studied the RR Lyrae star properties as a function of galactocentric radius in Phoenix and found no significant trends, consistent with the previously observed stellar population gradient for old stars in the dwarf galaxy. In particular, we found no significant trends in $\langle P_{ab}\rangle$, [Fe/H], or $A_V$ with respect to distance from the center of Phoenix in our RR Lyrae sample. We did however find a small but significant difference in mean RRc period, $\langle P_c\rangle$, between our two observed fields. Finally, we fit a closed-box chemical evolution model to the MDF of the RRab stars in Phoenix. Using the pulsation periods and amplitudes of the RRab stars, we calculated their metallicities using the period-amplitude-metallicity relation from \citet{alc00}. We obtained a mean metallicity for the probable RRab stars in Phoenix of $\langle [Fe/H]\rangle = -1.68 \pm 0.06$ dex. We found that a pure closed-box devoid of pre-enrichment poorly fit the MDF. However, a pre-enriched closed-box fits the MDF much better. The average metallicity associated with this best fit pre-enriched model, $[M/H] = -1.77 \pm 0.05$ dex, agrees well with the CEL for Phoenix derived by \citet{hid09}, supporting the notion that this galaxy chemically evolved similar to a pre-enriched, closed-box at a young age. Due to the isolated nature of Phoenix, we speculate that this pre-enrichment was not likely from any external source. Instead, we suggest that the prompt initial enrichment scenario in which material that contributed to the formation of Phoenix likely experienced an early generation of star formation marked by preferentially massive stars. | 14 | 4 | 1404.1037 |
1404 | 1404.3668_arXiv.txt | {% We study the behaviour of the van der Pol oscillator when either its damping parameter $\mu$ or its nonlinearity parameter $\xi$ is subject to additive or multiplicative random noise. Assuming various power law exponents for the relation between the oscillating variable and the sunspot number, for each case we map the parameter plane defined by the amplitude and the correlation time of the perturbation and mark the parameter regime where the sunspot number displays solar-like behaviour. Solar-like behaviour is defined here as a good correlation between the rise rate and cycle amplitude {\it and} the lack of a good correlation between the decay rate and amplitude, together with significant ($\ga 10$\,\%) r.m.s. variation in cycle lengths and cycle amplitudes. It is found that perturbing $\mu$ alone the perturbed van der Pol oscillator does not show solar-like behaviour. When the perturbed variable is $\xi$, solar-like behaviour is displayed for perturbations with a correlation time of about 3--4 years and significant amplitude. Such studies may provide useful constraints on solar dynamo models and their parameters. } | A complete spatial truncation of the nonlinear $\alpha\Omega$ dynamo equations with a dimensional analysis of some of the terms is known to give rise to a nonlinear oscillator equation of the form \begin{equation} \label{nemlinoegy} \ddot B = -\omega^2 B - \mu(\xi B^2-1)\dot{B} -\gamma B^3 \end{equation} where $B$ is the amplitude of the toroidal magnetic field, and the parameters $\mu$, $\xi$ and $\gamma$ may be expressed by the dynamo parameters (dynamo number, meridional flow amplitude, nonlinearity parameters) \citep{mininni2001,lopes2011}. This is a combination of the van der Pol and Duffing oscillators, the two most widely studied nonlinear oscillator problems. In the past decade, the possibility of representing the sunspot number series with such a nonlinear oscillator model was explored in a number of papers (e.g. \citealt{mininni2000}; \citealt{pontieri2003}; \citealt{lopes2009meridional}; \citealt{passos2012}). These studies demonstrated that with a suitable choice of parameters, the overall phase space structure of the sunspot number series can be well reproduced and cycle-to-cycle variations may also be qualitatively well modeled by admitting a stochastic perturbation to one or more of the model parameters. A more quantitative study of this problem should, however, also examine whether the behaviour of a stochastically perturbed oscillator of the form (\ref{nemlinoegy}) adheres to the known regularities in the cycle-to-cycle variation of solar activity. The most important such regularity is the Waldmeier effect. In its currently adopted formulation the effect consists in a good correlation (coefficient $r\simeq 0.85$) between the rise rate of a cycle and its maximum amplitude (\citealt{lantos}; \citealt{cameron}). On the other hand the {\it lack} of a statistically significant correlation between the decay rate and the cycle amplitude forms an equally important quantitative constraint. Taking into account the length of the sunspot number series this implies that $|r|\la 0.5$ for any such correlation. In order to study this problem we have started a systematic investigation of the parameter space of stochastically perturbed nonlinear oscillators of the type (\ref{nemlinoegy}). As a first step, here we consider a pure van der Pol oscillator, neglecting the last term in equation (\ref{nemlinoegy}). In the next section we present our stochastically perturbed van der Pol oscillator model in detail while results of the Monte Carlo simulations are presented and discussed in Section \ref{results}. Section 4 concludes the paper. | The present study is only the first step in a more extensive systematic investigation of the parameter space of the stochastically perturbed nonlinear oscillators potentially representing the sunspot number series. Here we focused on the behaviour of stochastically perturbed van der Pol oscillator when only one of its parameters (either the damping $\mu$ or the nonlinearity $\xi$) is varied. We focused on reproducing the Waldmeier effect and other known features of the solar cycle. It was found that in certain parameter ranges the stochastically perturbed van der Pol oscillator does display all the expected characteristics simultaneously but only when the perturbed oscillator parameter is $\xi$. Another interesting conclusion from these studies is that a good correlation between cycle amplitude and rise rate (Waldmeier effect) is displayed by a fairly wide class of nonlinear oscillators; at the same time the {\it lack} of a similar good correlation between cycle amplitudes and decay rates is a potentially even more stringent condition for solar-like behaviour that is in fact often harder to reproduce than the Waldmeier effect itself. Our finding that random perturbations of $\mu$ alone do not lead to solar-like behaviour is of some interest as in at least one earlier study (\citealt{lopes2008}) cycle variations in the observed SSN series were interpreted in terms of variations in the speed of meridional circulation which, in the truncated flux transport dynamo considered there, mainly influence the oscillator through the damping parameter. This underlines that while fitting the observed cycles with a nonlinear oscillator with varying parameters results in parameter variations that may seem to be random at first sight, the random nature of such variations is actually not demonstrated and a systematic study of an ensemble of oscillators with truly random parameter variations is needed to confirm or discard solar-like behaviour. It should be stressed, however, that this conclusion is still of a preliminary nature and it may change if e.g. the perturbations in $\mu$ and $\xi$ are assumed to be interrelated, as expected for a flux transport dynamo. Possibilities for future extensions of this study therefore include a joint perturbation of $\mu$ and $\xi$ and the inclusion of a (possibly also perturbed) third order term in the oscillator equation, as in equation (1). | 14 | 4 | 1404.3668 |
1404 | 1404.4879_arXiv.txt | We present multi-wavelength observations that trace more than 40 years in the life of the active galactic nucleus (AGN) in Mrk\,590, traditionally known as a classic Seyfert 1 galaxy. From spectra recently obtained from \HST, {\it Chandra}, and the Large Binocular Telescope, we find that the activity in the nucleus of Mrk\,590 has diminished so significantly that the continuum luminosity is a factor of 100 lower than the peak luminosity probed by our long baseline observations. Furthermore, the broad emission lines, once prominent in the UV/optical spectrum, have all but disappeared. Since AGN type is defined by the presence of broad emission lines in the optical spectrum, our observations demonstrate that Mrk\,590 has now become a ``changing look'' AGN. If classified by recent optical spectra, Mrk\,590 would be a Seyfert $\sim$1.9$-$2, where the only broad emission line still visible in the optical spectrum is a weak component of \Halpha. As an additional consequence of this change, we have definitively detected UV narrow-line components in a Type 1 AGN, allowing an analysis of these emission-line components with high-resolution COS spectra. These observations challenge the historical paradigm that AGN type is only a consequence of the line of sight viewing angle toward the nucleus in the presence of a geometrically-flattened, obscuring medium (i.e., the torus). Our data instead suggest that the current state of Mrk\, 590 is a consequence of the change in luminosity, which implies the black hole accretion rate has significantly decreased. | The dichotomy separating active galactic nuclei (AGN) into Type 1 --- those observed to have broad emission lines --- and Type 2 --- those without broad lines --- originated from the first observations of \citet{Seyfert43} that demonstrated that the nebular lines in the nuclei of these nearby ``Seyfert'' galaxies sometimes had broad emission wings superposed with a narrow core and sometimes did not. A working definition of Type 1 and Type 2, based on the relative line widths of the forbidden and Balmer lines was then elucidated by the work of \citet{Khachikian&Weedman74}. \citet{Osterbrock&Koski76} and \citet{Osterbrock77, Osterbrock81} later note that the observed emission spectra of Seyfert galaxies are not so simple and argue for a continuum of intermediate types between these two extremes, e.g., Seyfert 1.2, 1.5, 1.8, and 1.9 based on the relative strength of the narrow \Hbeta\ component with respect to the broad component. \citet{Antonucci85} presented a plausible scenario to physically explain this dichotomy by invoking different line of sight orientations: the broad line-emitting region (BLR) in Type 2 AGN (Sy2s) is obscured by an optically thick, geometrically flattened medium (often referred to as the torus) leading to a ``hidden" BLR (HBLR), whereas the BLR is visible in Type 1 AGN (Sy1s) because our line of sight is not obscured by this optically thick medium. This unified model of AGN \citep[see also][]{Antonucci93} followed from the discovery by \citet{Antonucci85} that broad emission lines were present in the polarization spectrum of NGC\,1068. The key assumption in this model is that all AGN systems are physically similar, with generally disk-like geometries. Thus, it is only the line of sight orientation of the BLR with respect to the obscuring medium that results in the observed type differences. The unified model for Sy1 and Sy2 systems has been challenged by observations that suggest that not {\it all} AGN conform to this paradigm. Sy2s that do not have broad lines observed even through spectropolarimetry have been coined non-HBLR or ``true" Sy2s, and some may, in fact, be ``bare" Sy2s that are postulated to be devoid of the typical obscuring medium surrounding the BLR \citep[see e.g.,][]{Barth99, Tran01, Tran03, Panessa&Bassani02, Laor03, Zhang&Wang06}. On the other hand, \citet{Antonucci12} advises caution before classifying objects as non-HBLR objects, as finding reflected broad lines is largely serendipitous and dependent on the particular source having a well-placed scattering medium with the right properties. Thus, the existence of true or bare Sy2's may not be well quantified. Nonetheless, there are also objects known as ``changing look" AGN, which are more easily identifiable because an obvious change has occurred in the observed spectrum to warrant the application of this classification. This characterization was originally coined based on X-ray observations in which sources appear alternately Compton-thin or ``reflection-dominated" (likely Compton thick) over the course of years \citep[e.g.,][]{Bianchi05}. This qualifier has recently been extended to include objects which sometimes appear to have Sy2 and sometimes Sy1 characteristics in their optical spectrum \citep[e.g.,][]{Shappee13}. Several examples of these optical changing look AGNs appear the literature over the years. One of the most well-cited early examples is that of NGC\,4151, one of Seyfert's original galaxies assigned Type 1.5 by \citet{Osterbrock77}, but in which the optical broad lines all but disappeared (except for weak and possibly asymmetric wings) in the 1980's \citep{Antonucci83, Lyutyj84, Penston84} and have since returned \citep[see, e.g.,][]{Shapovalova10}. There are two typically accepted postulates to explain AGN changing their type: (1) variable obscuration or (2) variable accretion rate. Variations in the obscuring medium is more suited to the unification paradigm of type resulting from different viewing angles and is typically invoked to explain changing look AGN in the X-ray regime \citep[e.g.,][]{Bianchi05, Risaliti09, Marchese12, Marin13}. In contrast, it has been suggested that variations in luminosity and accretion rate are not only responsible for type changes in individual objects but are also responsible for the whole AGN typing sequence. In other words, the structure of the BLR changes with accretion rate, and objects evolve from high accretion rate when the AGN turns on --- Type 1 --- to low accretion rate once the AGN has depleted its nuclear material --- Type 2 \citep{Tran03, Wang&Zhang07, Elitzur14}, possibly oscillating between Type 1 and Type 2 and/or intermediate types between high and low accretion states while the nucleus remains active \citep{Penston84, Korista&Goad04}. At sufficiently low accretion rates, many postulates exist demonstrating that a radiatively efficient BLR simply cannot be supported, due to, e.g., (1) a dearth of ionizing photon flux \citep{Korista&Goad04}, (2) the critical radius at which the accretion disk changes from gas pressure-dominated to radiation pressure-dominated becoming smaller than the inner-most stable orbit \citep{Nicastro00, Nicastro03}, (3) mass conservation considerations in the paradigm that the BLR arises in a disk wind with a fixed radial column, where the mass outflow rate cannot be sustained \citep{Elitzur&Ho09}, or (4) the accretion disk structure changing, replacing the disk-wind BLR with a radiatively inefficient accretion flow (RIAF) consisting of a fully ionized, low-density plasma incapable of producing broad lines \citep{Trump11}. \citet{Laor03} similarly suggests a bolometric luminosity below which the BLR cannot exist, though with a slightly different argument following from a maximally observed broad line width. Observational results suggest that both of these physical processes are likely at play in different objects. For example, \citet{Alexander13} present a serendipitously detected {\it NuSTAR} source at redshift $z=0.510$ for which observations suggest that significant obscuration may have moved into our line of sight to the nucleus of this source, changing it from a Type 1 to Type 2: SED template fitting \citep[following][]{Assef10} show its SED is currently consistent with other optically identified Type 2 AGN in the same sample and an $E(B-V) = 0.6 \pm$0.5 mag, yet based on archival observations, its SED was previously consistent with that of a Type 1 AGN with estimated $E(B-V) = 0.00\pm$0.01 mag. \citet{Alexander13} also presents a recent optical spectrum of this source, J183443+3237.8, that shows narrow emission lines and significant reddening of the continuum, though there is evidence for a weak broad component of \mgii. On the other hand, \citet{Shappee13atel} report on the interesting behavior of NGC\,2617, which has changed from a Sy1.8 \citep{Moran96} to a Sy1, likely (though not conclusively) due to a recent outburst that triggered a transient source alert within the All-Sky Automated Survey for SuperNovae (ASAS-SN\footnote{http://www.astronomy.ohio-state.edu/$\sim$assassin}). Subsequent observations of this AGN demonstrated a continued increase of the optical through X-ray emission by about an order of magnitude compared to past observations \citet{Shappee13}. Due to the outburst nature of this activity, which occurred over relatively short time scales, it is highly unlikely that obscuration moving out of the line of sight could be responsible for this change in type. Additionally, the optical spectral changes between Type 2 and Type 1.9 of NGC\,2992 also seem to be at least loosely correlated with large variability amplitudes (factors of a few tens) observed in the X-rays over both short (year) and long (decades) time scales, such that broad \Halpha\ only seems to be observed coincident with high X-ray states \citep[see][]{Gilli00, Murphy07, Trippe08}. In this work, we examine the interesting and extreme phenomenon of Mrk\,590 (alt.\ NGC\,863) --- a classic Sy1 galaxy \citep{Osterbrock77, Weedman77} at $z=0.026385$ --- which appears to have changed from Type 1.5 to Type 1 then to Type $\sim$1.9$-$2, based on chronologically catalogued multi-wavelength observations. Observations from MDM Observatory in late 2012 revealed that the broad Balmer emission lines seem to have disappeared from Mrk\,590. Observations with higher S/N obtained in 2013 February with the MODS1 spectrograph on the Large Binocular Telescope confirmed that there was no longer any broad component to \Hbeta\ and possibly only a very weak broad \Halpha\ component. Because this system is so well-studied, we have been able to gain additional understanding of this change through multi-wavelength observations taken in its past and present states. In Section \ref{S_data} we describe the new and archival data we have gathered. We present our optical continuum fitting methods in Section \ref{S_hostfit} and discuss the overall observed trends in Section \ref{S_40yearprop}. Section \ref{S_discuss} follows with a discussion of the properties of the nuclear region of Mrk\,590 as well as potential implications of the observed behavior of Mrk\,590. We summarize our findings and plans for future work in Section \ref{S_summary}. A cosmology with $\Omega_{m}=0.3$, $\Omega_{\Lambda}=0.70$, and $H_0 = 70$ km sec$^{-1}$ Mpc$^{-1}$ is assumed where necessary. | \label{S_discuss} The seeming ability of objects like Mrk\,590 to transition between Type 1 and Type 2 without an otherwise obvious flaring event like what \citet{Shappee13} observed in NGC\,2617 raises the question of what assigning a type is really saying about an AGN. It is possible that such classifications should be taken as more of a statement of current `state' rather than a true, fixed `type', as Mrk\, 590 further indicates that type can change. This distinction may be of particular importance since it is becoming apparent that whether a particular AGN appears at any given time to be Type 1 or Type 2 (or somewhere in between) may be saying something more significant about the AGN physics and the immediate environment of the black hole at that particular time than strictly about the viewing angle. For Mrk\,590, the decrease in the BLR, the UV, and the X-ray emission all seem to point to a decrease in total luminosity, thus accretion rate, and not simply a change in obscuration. Furthermore, the dynamical timescale of the \Hbeta\ emitting region is roughly 8 yrs in the rest frame of Mrk\,590, based on the black hole mass and \Hbeta\ FWHM given by \citet{Peterson04}. An optically thick medium capable of occulting the BLR could only reside outside the dust sublimation radius and would therefore have a much longer dynamical timescale. Yet, the broad \Hbeta\ emission transformed from strong to non-existent in $\lesssim$10 years (1996--2006). Furthermore, we see a dramatic change in the strength of not only the BLR emission but also the continuum and the NLR emission. If obscuration were responsible for this, the obscuring medium would need to cover both our line of sight to the continuum source, in which case we may expect to see evidence of an obscurer in our X-ray continuum model, which we do not, as well as that between the continuum and the NLR, which is highly improbable. Therefore, though it cannot be definitively ruled out, line of sight obscuration is an unlikely explanation for the observations we present here. In general, the acquisition of sufficient data to determine the bolometric luminosity and spectral energy distribution of an AGN before and after a change in type should be able to distinguish between these two physical origins for the change. Changes in accretion rate should affect the flux at all wavelengths. However, flux changes due to obscuration will be wavelength dependent, where the decrease in UV/optical emission will be countered by an increase in the mid-to-far IR flux, since the obscuring medium absorbs the higher energy photons but then reradiates them as thermal emission at these long wavelengths. A more likely explanation for the suite of observations presented here is that an accretion event occurred 40$+$ years ago. This event was capable of stimulating the production of ionizing photons that have excited the BLR emission over the subsequent 40 years. However, the energy produced in such an event has now been depleted. The symmetry of the narrow emission lines and the time scales over which we see significant changes in both the broad and narrow emission lines are too short for it to be plausible that this accretion event actually created these regions, i.e., to initially form and light up the AGN from a quiescent state. Instead, it is likely that at least the bulk of the nuclear gas was already present, just not emitting significant broad-line emission. This suggests that the appearance of a BLR may be episodic over the accretion history of supermassive black holes, depending on the reservoir of material available in the nucleus for accretion. It further demonstrates the importance of repeat observations of larger samples of Seyfert galaxies over longer baselines as a means to better quantify how often this behavior may occur. Transition objects like Mrk\,590 may also be a link to other `abnormal' AGN found from large survey samples. For example, \citet{Roig14} found unusual broad-line objects among the BOSS sample of luminous galaxies that show a broad \mgii\ emission line component but little to no broad Balmer emission. An interesting possibility is that these objects are in a similar transition state as Mrk\,590. Unfortunately, a recent spectrum of the \mgii\ region in Mrk\,590 has not yet been obtained, so more data are needed to connect the likely time variable nature of the broad Balmer lines in Mrk\,590 with this larger class of objects. We have been awarded time to obtain a spectrum of the \mgii\ region in Mrk\,590 with \HST\ in Cycle 21 for this purpose and will therefore be able to address this in future work. Under the assumption that the observed behavior of Mrk\,590 {\it was} the consequence of a past accretion event, we can infer additional physical properties of this system. The BLR would have known and responded to a new accretion event almost instantaneously, given the \Hbeta\ radius of $\sim$10$-$20 light days \citep{Peterson04} and the short recombination time in this region. However, this is not the case for gas at the distance and densities of the NLR. The behavior and subsequent properties of the \ob\ lines, as well as the heterogeneous data in our sample, allow us to put various independent limits on the NLR radius. One of the weakest, but nonetheless most reliable constraints on the NLR size comes from the fact that we measure nearly identical [\oiii] fluxes from the 2003 SDSS spectrum as we do from the large aperture RM campaign spectra taken in the 1990's. This indicates that most of the [\oiii] emission must arise from the spatially unresolved region with a 1.7\,kpc diameter --- the physical size subtended by the SDSS aperture at the distance of Mrk\,590. Next, insofar as we can rely on the observed increase of the [\oiii]\,$\lambda$5007 flux by 40\% between 1973 and 1983, light travel time arguments suggest that the NLR emission must be $\lesssim3$\,pc from the central source. In addition, the fluxes measured from the 2003 SDSS spectrum show that the continuum and BLR emission had significantly declined already, yet the [\oiii] flux has not yet changed. However, the [OIII] flux {\it had} diminished by the time the 2006 MDM spectrum was obtained (again assuming at least some amount of trust in this flux measurement). Thus, even if this decline began right after the final RM campaign [OIII] flux calibration spectrum was obtained, we deduce a light travel timescale of $\lesssim$10 years, consistent with the distance inferred from the earlier increase in flux. Next, we briefly consider the velocities of the [\oiii]-emitting gas, though the evidence for such a discussion is less robust than that derived from the fluxes. This is because large ($\sim$100 \kms) systematic errors in the velocity widths derived from individual spectra taken through large spectroscopic apertures are possible. This is on account of the difficulty involved in accurately measuring and correcting for the spectral resolution given that the PSF does not fill the slit. \citet{Vrtilek&Carleton85} measure the [\oiii]\,$\lambda$5007 velocity width to be 397$\pm$22\,\kms\ from a high (23\,\kms) resolution spectrum taken sometime between 1980 June and 1981 July. This is consistent with our resolution corrected velocities, for the most part. If we at least assume that these velocities approximately represent the virial velocity at the radius of the NLR, this also indicates consistent distances with other limits we considered above, with velocities $\sim$250$-$550\,km s$^{-1}$ corresponding to distances of $\sim$0.7$-$3.3\,pc. While we regrettably cannot trust the reliability of the 1973 line width measurement, it is notably larger than the other measurements. This could lead to the interesting, though highly speculative, consideration that with this early observation, we are actually observing the NLR `lighting up' from the inside out as the AGN continuum luminosity begins to increase significantly. Assuming a 10$^4$ K gas, the recombination time for gas near the critical density of \ob\ will be on the order of days; however, as the density drops with increasing radius to the expected average density of the NLR of $\sim$2000 cm$^{-3}$ \citep{Koski78}, the recombination time approaches a decade. Such effects could lead to a scenario in which we see narrow-line emission only from the higher-velocity, more dense gas in the innermost regions of the NLR as the AGN continuum first brightens, and only after years will the velocity widths of the integrated [\oiii]-emitting gas decrease as the spatial extent of the emission increases. Certainly, the range of velocities measured from the narrow component of the various emission lines we consider, which are significantly larger than the line width uncertainties, indicate that ``narrow'' emission line gas is emitted from an extended spatial region with a density gradient. The innermost emission from the high ionization species, such as \civ, \Lya, and \heii\ is coming from well within sub-pc scales. Finally, to complete the picture of events, our observations indicate that as the AGN continuum luminosity increased (decreased), the BLR emission strengthened (diminished) faster than the NLR emission. This behavior suggests the spatial propagation of information from the central source outward because of density effects and light travel time. This again implies that an ``event" of some sort may have triggered the creation of enough ionizing photons to excite the observed changes in BLR emission and subsequent strengthening of the NLR emission (as discussed above). There is nothing in the observations that can yet point to ``what" may have been accreted to cause the observed sequence of events. However, we can {\it roughly} estimate the total mass needed to power the bolometric luminosity integrated over the span of our observations with a simple, back of the envelope calculation. We take the 5100\,\AA\ fluxes from all optical spectra shown in Figure~\ref{fig_allspec} (values in Table 1) and convert these to bolometric luminosities assuming a bolometric correction of 8.10 \citep{Runnoe12err}. We then estimate the integrated luminosity both under the rough continuum emission light curve shown in Figure \ref{fig_lightcurves} and by extending the observed rise and decline, modeled as Gaussian wings, to both the past and future to conservatively cover a temporal range from 1950 to 2020. Finally, assuming that the mass-to-energy conversion efficiency is 7\%, we find that the observed continuum output during this event can be accounted for with only $\sim$1$-$2\,$M_{\odot}$ of total mass being accreted over these 70 years. We have presented optical, UV, and X-ray observations of the `Classical' Seyfert 1 Mrk\,590 that span the past 40+ years. This interesting object brightened by a factor of 10's between the 1970's and 1990's and then faded by a factor of a 100 or more at all continuum wavelengths between the mid-1990's and present day. Notably, there is no evidence in the current data set that this recent, significant decline in flux is due to obscuration; in particular, the most recent X-ray observations are consistent with zero intrinsic absorption. There were similarly dramatic changes in the emission-line fluxes. The most striking change is the complete disappearance of the broad component of the \Hbeta\ emission line, which had previously been strong (equivalent widths $\sim$20$-$60 \AA) and the focus of a successful reverberation mapping campaign that resulted in a secure estimate of the supermassive black hole contained within the nucleus of Mrk\,590 of $\sim5\times10^7 M_{\odot}$. As a result of these significant changes, the optical spectrum of Mrk\,590 currently looks more like that of a Sy2 AGN, with predominantly only narrow emission lines and a strong host galaxy stellar continuum. The changes in emission line properties over this time period allowed us to determine that Mrk\,590, at least in its current state, has NLR emission line ratios that are similar to those measured in Sy2 spectra. We were also able to put limits on the radius of the [\oiii]-emitting region of the NLR to be $\sim$0.7$-$3.3\,pc based on the observed changes in the integrated line flux and the velocity width of [\oiii]\,$\lambda$5007. These results are consistent with the NLR size determined by \citet{Peterson13} for NGC\,5548, another nearby Sy1 galaxy containing a BH of similar mass to Mrk\,590. The implications arising from this long time series of Mrk\, 590 are that (1) Mrk\,590 is a direct challenge to the historical paradigm that AGN type is exclusively a geometrical effect, and (2) there may not be a strict, one-way evolution from Type 1 to Type 1.5$-$1.9 to Type 2 as recently suggested by \citet{Elitzur14}. Instead, for at least some objects, the presence of BLR emission may coincide only with episodic accretion events throughout a single active phase of an AGN. If true, such behavior may be more prominent in Seyfert galaxies, where accretion has been theorized to be a consequence of secular processes and therefore likely more episodic than quasar activity, which may be triggered more predominantly by major mergers \citep[see, e.g.,][]{Sanders88, Treister12}. The final possibility, of course, is that we are witnessing the final stages in the life of this AGN, and Mrk\,590 is completely turning off. This would be an incredible find, but is likely the most improbable explanation, given the low duty cycles for low redshift AGN \citep[see, e.g.,][]{Schulze&Wisotzki10, Shankar13}. We plan to continue monitoring Mrk\,590 to look for further changes in its current state, but only time and more observations will tell whether the BLR returns. | 14 | 4 | 1404.4879 |
1404 | 1404.6090_arXiv.txt | We used the SPIRE/FTS instrument aboard the Herschel Space Observatory (HSO) to obtain the Spectral Line Energy Distributions (SLEDs) of CO from J=4--3 to J=13--12 of Arp\,193 and NGC\,6240, two classical merger/starbursts selected from our molecular line survey of local Luminous Infrared Galaxies (LIRGs: $\rm L_{IR}$$\geq $10$^{11}$\,L$_{\odot}$). The high-J CO SLEDs are then combined with ground-based low-J CO, $^{13}$CO, HCN, HCO$^{+}$, CS line data and used to probe the thermal and dynamical states of their large molecular gas reservoirs. We find the two CO SLEDs strongly diverging from J=4--3 onwards, with NGC\,6240 having a much higher CO line excitation than Arp\,193, despite their similar low-J CO SLEDs and $\rm L_{FIR}/L_{CO,1-0}$, $\rm L_{HCN}/L_{CO}$ (J=1--0) ratios (proxies of star formation efficiency and dense gas mass fraction). In Arp\,193, one of the three most extreme starbursts in the local Universe, the molecular SLEDs indicate a small amount ($\sim $5\%-15\%) of dense gas (n$\geq$$10^{4}$\,cm$^{-3}$) unlike NGC\,6240 where most of the molecular gas ($\sim $60\%-70\%) is dense (n$\sim $($10^4$--$10^5$)\,cm$^{-3}$). Strong star-formation feedback can drive this disparity in their dense gas mass fractions, and also induce extreme thermal and dynamical states for the molecular gas. In NGC\,6240, and to a lesser degree in Arp\,193, we find large molecular gas masses whose thermal states cannot be maintained by FUV photons from Photon Dominated Regions (PDRs). We argue that this may happen often in metal-rich merger/starbursts, strongly altering the initial conditions of star formation. ALMA can now directly probe these conditions across cosmic epoch, and even probe their deeply dust-enshrouded outcome, the stellar IMF averaged over galactic evolution. | The discovery of bright CO J=1--0 line emission in the Orion nebula (Wilson, Jefferts, \& Penzias 1970) opened up the rich field of molecular astrophysics with molecular lines as the primary probes of the physical conditions of Giant Molecular Clouds (GMCs), the most massive structures in galaxies and the sites of star formation. The much weaker rotational transitions from high-dipole moment molecules of CS and CN, which probe much denser gas, were detected soon afterwards (Wilson et al. 1971). As receiver sensitivities improved multi-J transitions of such high-dipole molecules (mostly HCN and HCO$^{+}$) were used along with those of CO to probe the full range of physical conditions of the molecular gas in nearby star-forming (SF) galaxies (e.g. Solomon et al 1992; Gao \& Solomon 2004; Gracia-Carpio et al. 2007; Krips et al. 2008; Greve et al. 2009). These observational studies, and theoretical investigations of the supersonic turbulent GMCs either as individual entities (Li et al. 2003; Larson 2005; Jappsen et al. 2005) or embedded within galaxies (Krumholz \& McKee 2005), showed the dense molecular gas (n$\ga $10$^{4}$\,cm$^{-3}$) as the phase where stars form. Its physical conditions are thus the crucial input for all star formation theories and the resulting stellar Initial Mass Function (IMF) (Larson 2005; Elmegreen et al. 2008). However, the weakness of high-dipole moment molecular lines (e.g. HCN J=1--0 is $\sim $5-100 times fainter than CO J=1--0) that trace high density gas prevented large extragalactic surveys of such lines, while strong atmospheric absorption limits observations of the more luminous CO SLEDs mostly up to J=3--2 (e.g. Yao et al. 2003; Leech et al. 2010), i.e. the first CO transition that starts tracing solely the dense and warm SF gas ($\rm n_{crit}$(3-2)$\sim $10$^{4}$\,cm$^{-3}$, $\rm E_{3}/k_B$$\sim $33\,K). Such low-J CO line spectroscopy has no diagnostic value regarding the conditions of the dense gas, and little overlap with the CO SLEDs of distant SF galaxies where CO J=3--2, 4--3 and higher-J transitions are mostly detected, redshifted into more transparent mm/submm atmospheric windows (e.g. Solomon \& Vanden Bout 2005; Weiss et al. 2007). The importance of high-J CO lines in probing the dense gas physical conditions was recently underscored by a small extension of CO SLEDs to include J=4--3 and 6--5 for a few local Luminous Infrared Galaxies (LIRGs: $\rm L_{IR}$$\geq $$10^{11}$\,L$_{\odot}$). These revealed large dense {\it and} warm gas reservoirs that are irreducible to ensembles of Photon-Dominated Regions (PDRs) in some merger/starbursts (Papadopoulos et al. 2012a). Such conditions were also found in Mrk\,231 and Arp\,220 using CO SLEDs from J=1--0 up to J=13--12 obtained with SPIRE/FTS and ground-based observations (van der Werf et al. 2010; Rangwala et al. 2011). Finally high-J CO and heavy rotor molecular lines are necessary for better estimates of the $\rm X_{CO}$=$\rm M_{tot}(H_2)/L_{CO,1-0}$ factor in merger/starbursts where, unlike isolated spirals, the dense phase can contain much of their total molecular gas mass (Papadopoulos et al.~2012b). The thermal, dynamical and chemical state of the dense gas in SF galaxies, its mass contribution to $\rm M_{tot}$(H$_2$), the effects of SF and AGN feedback, and complete CO SLEDs from J=1--0 up to high-J transitions as local benchmarks for high-z CO observations were the key drivers for our Herschel Comprehensive (U)LIRG Emission Survey (HerCULES), an open time Key program (PI: Paul van der Werf) on the ESA Herschel Space Observatory (HSO)\footnote{Herschel is an ESA space observatory with science instruments provided by European-led Principal Investigator consortia and with important participation from NASA} (Pilbratt et al. 2010), augmented by a large ground-based low-J CO and $^{13}$CO line survey (Papadopoulos et al. 2012a). Here we report on HSO SPIRE/FTS and ground-based observations of Arp\,193 and NGC\,6240, two prominent merger/starbursts from HerCULES whose similar low-J CO SLEDs, $\rm \epsilon_{SF, co} $=$\rm L_{FIR}/L_{co, 1-0}$ (a proxy of SF efficiency SFE=SFR/$\rm M_{tot}(H_2)$) and $\rm r_{HCN/CO}$=$\rm L^{'}_{\rm HCN,1-0}$/$\rm L^{'} _{\rm CO,1-0}$ (a proxy of $\rm f_{dense}$=M(n$>$10$^{4}$\,cm$^{-3}$)/M$_{\rm tot}$(H$_2$)) ratios, make them good testbeds for exploring diferences in their dense gas properties. This work is structured as follows: 1) we present the SPIRE/FTS and ground-based molecular line data and their reduction (Section 2), 2) we construct the full CO SLEDs from J=1--0 to J=13--12 and use them along with our $^{13}$CO, HCN, HCO$^{+}$, and CS line data to find the average conditions and mass of the molecular gas components using radiative transfer models (Section 3), 3) determine their thermal states and energy requirements (Section 4), and 4) discuss general implications for the ISM in merger/starbursts, and present our conclusions (Section 5). We adopt a flat $\Lambda $-dominated cosmology with $\rm H_0$=71\,km\,s$^{-1}$\,Mpc$^{-1}$ and~$\Omega_{\rm m}$=0.27. | 14 | 4 | 1404.6090 |
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1404 | 1404.3817_arXiv.txt | The detection of B-mode shows a very powerful constraint to theoretical inflation models through the measurement of the tensor-to-scalar ratio $r$. Higgs boson is the most likely candidate of the inflaton field. But usually, Higgs inflation models predict a small value of $r$, which is not quite consistent with the recent results from BICEP2. In this paper, we explored whether a cosmological constant energy component is needed to improve the situation. And we found the answer is yes. For the so-called Higgs chaotic inflation model with a quadratic potential, it predicts $r\approx 0.2$, $n_s\approx0.96$ with e-folds number $N\approx 56$, which is large enough to overcome the problems such as the horizon problem in the Big Bang cosmology. The required energy scale of the cosmological constant is roughly $\Lambda \sim (10^{14} \text{GeV})^2 $, which means a mechanism is still needed to solve the fine-tuning problem in the later time evolution of the universe, e.g. by introducing some dark energy component. | Recently the detection of B-mode from CMB by the BICEP2 group \cite{Ade:2014xna} has indicated a strong evidence of inflation \cite{Guth:1980zm,Linde:1981mu,Albrecht:1982wi}, which solves many theoretical puzzles in the Big Bang cosmology. The B-mode polarization can be only generated by the tensor perturbations. According to the reports of the BICEP2 experiment, the tensor-to-scalar ratio is in range: $r=0.20_{-0.05}^{+0.07} (68\%$ CL). In a simplest slow-roll inflation model, the early universe was driven by a single scalar field $\phi$ with a very flat potential $V(\phi)$. Usually, we call this field the inflaton. Although there are many inflation models in the market, we still do not well-understand what is the inflaton. The most economical and fundamental candidate for the inflaton is the standard model (SM) Higgs boson, which has been already observed by the collider experiment LHC in 2012 \cite{Aad:2012tfa,Chatrchyan:2012ufa}. In this sense, Higgs inflation is a simple and elegant model. However, it is not easy for the Higgs boson to realize a inflation model with correct density perturbations. To see this, we estimate the inflaton mass from the amplitude $A_s$ of the scalar perturbation power spectrum in the chaotic inflation model \cite{Linde:1983gd} with a quadratic potential $V(\phi)=m^2\dot\phi^2/2$: \begin{equation}\label{equ:mass1} m \approx 1.5\times 10^{13}\left(\frac{N}{60}\right)^{-1}\left( \frac{ 10^{9}A_s }{2.19}\right)^{1/2} \text{GeV} \,, \end{equation} which is many orders of magnitude larger than the observed Higgs mass, $m_{h}\approx 125.9\pm0.4$ GeV. In other words, the potential of Higgs field $h$ is not flat enough to realize an inflation. By introducing a non-minimal coupling to the gravity ($\sim h^2 R$) , one could indeed achieve such a flat potential \cite{Bezrukov:2007ep} after a conformal transformation. And the predictions of this kind of non-minimal coupling Higgs inflation are well consistent with observations before BICEP2. The authors in ref.\cite{Cook:2014dga} have found that this model can not accommodate the new measurement from BICEP2, because it generally predicts a small amplitude of tensor perturbations. An alternative Higgs inflation model was proposed in ref.\cite{Germani:2010gm}, in which the Higgs boson kinetic term is non-minimally coupled to the Einstein tensor ($\sim G^{ab}\partial_a h\partial_b h$). According to the recent analysis on this model \cite{Germani:2014hqa}, it predicts $r\approx0.16$ when the number of e-folds $N\approx33$, since $r\approx16/(3N+1)$ in this model. However, to overcome the problems in the Big Bang theory, the number of e-folds is required to be around $N\approx60$, then the tensor-to-scalar ratio becomes even smaller, say $r\approx 0.09$. Another interesting Higgs inflation model called the Higgs chaotic inflation is proposed in ref.\cite{Nakayama:2010sk}. In this model, the SM Higgs boson realizes the quadratic chaotic inflation model, based on the so-called running kinetic inflation \cite{Nakayama:2010kt, Takahashi:2010ky}. The kinetic term of the inflaton is significantly modified at large field values, while it becomes the canonical one when $h$ is small. The value of $r$ in this model is the same as that in the chaotic inflation model with a quadratic potential, i.e. $r=8/N$. For $N\approx 60$, it predicts $r \approx 0.13$, but if we require a larger $r$, say $r\approx 0.2$, a smaller $N$ is needed, say $N \approx 40$, which is a little better than that predicted in the other Higgs inflation models, see ref.\cite{Nakayama:2014koa} for recent revisited in this model. It seems that the Higgs chaotic inflation is a charming Higgs inflation model in the market. On the other hand, there is a challenge for a single field inflation with BICEP2 result. For the chaotic inflation, the larger the value of the tensor-to-scalar ratio is, the smaller the value of the running of the spectral index is, see the details in ref.\cite{Gong:2014cqa}. Therefore, to be more consistent with observations, one might consider a little more beyond a single inflation model. Among many choices, the cosmological constant is often forgotten when one building an inflation model, since by itself only the exact scale-invariant Harrizon-Zel'dovish power spectrum with the scalar spectral index $n_s =1$ could be produced, which is already ruled out at over $5\sigma$ by \textit{Planck} \cite{Ade:2013uln}. However, we find that the situation is changed when the early universe is dominated by the cosmological constant as well as the inflaton. It could give $n_s \approx 0.96$, $r\approx 0.2$ when the number of e-folds is not so small, say $N\approx 56$, and it also predict the correct magnitude of the spectrum amplitude. In the following, we will assume that the running kinetic approach is a correct way to realize inflation by SM Higgs boson and we also assume that both inflaton and the cosmological constant dominated the universe during the inflation time. In next section, we give a briefly review of the running kinetic inflation and then we pursue the role played by the cosmological constant during inflation. Finally, we will draw our conclusions and give some discussions in the last section. | The recent detection of B-mode by BICEP2 indicates an exciting leap forward in our ability to explore the early universe and fundamental physics. The measurement of the tensor-to-scalar ratio $r\approx0.2$ shows a very powerful constraint to theoretical inflation models. Higgs boson is the most likely candidate of the inflaton field. However, its mass $m_h\sim\mathcal{O}(10^2)$ GeV is much smaller than that for a inflaton $m\sim\mathcal{O}(10^{13})$ GeV. To solve this hierarchy problem, a non-minimal coupling between the Higgs boson and gravity or a non-canonical kinetic term is needed. Usually, these Higgs inflation models predict a small value of $r$, which is not quite consistent with the results from BICEP2. In this paper, we explored whether a cosmological constant energy component is needed to improve the situation. And we found the answer is yes. The Higgs chaotic inflation now predicts $r\approx 0.2$, $n_s\approx0.96$ with e-folds number $N\approx 56$, which is large enough to overcome the problems in the Big Bang cosmology. However, we are still far from understanding the cosmological constant. And we haven't solve its fine-tuning problem in the later time evolution of the universe, which is asked why the present value of the cosmological constant is so small, or why the universe is accelerating at present $z\sim 1$. Noticed that the slow-roll parameters have a finite maximum value from Eqs.(\ref{equ:sl1}) and (\ref{equ:sl2}) as long as $\Lambda \neq 0$: $\epsilon_{\text{max}} \approx m^2/\Lambda$ when $\phi=\sqrt{2\Lambda}/m$, and $\eta_{\text{max}} \approx m^2/\Lambda$ when $\phi\rightarrow 0$. It seems that the inflation will never end if $\Lambda > m^2$. To end the inflation, one may need a phase transition of a heavy Higgs boson $\chi$ with its mass at GUT scale, and it also slightly couples to the light one that responsible to inflation by $\sim h^2\chi^2$. At the beginning of inflation, the heavy Higgs boson is stable at its true vacuum ($\chi=0$), then it only contributes a constant potential, which can be regarded as the cosmological constant. When the inflaton rolls down the potential and becomes small enough, the vacuum at $\chi=0$ turns to be a false one and the heavy boson would be no longer stable, then it rolls to true vacuum to end the inflation. In fact, the endless inflation is essentially due to the cosmological fine-tuning problem. Once a correct mechanism is found to reduce $\Lambda$ to its present observational value, then the inflation would be certainly end. We will give a concrete example in detail to realize such a mechanism that may solve the fine-tuning problem in the later work \cite{fengli}. The challenge for a single field inflation to predict a large value of the running of the index still exit, $n_s' \equiv dn_s/d\ln k \approx -0.00025$ for $r\approx0.2$ in our case, see also ref.\cite{Gong:2014cqa} for detail discussions on this issue. But the constraint on the running is not so tight: $n_s'\approx-0.013\pm 0.009 (68\%\text{CL})$ from the analysis of \textit{Planck} data, see ref.\cite{Ade:2013uln}. Furthermore, if additional sterile neutrino species are taken into account in the universe, one could also obtain $r\approx0.20$ without the running of the spectral index ($n_s'\sim0$), see refs.\cite{Zhang:2014dxk,Dvorkin:2014lea,zhang1404.3598}. Certainly, if a large running is well-confirmed in future, then other mechanisms explain it are urgently needed. | 14 | 4 | 1404.3817 |
1404 | 1404.1189.txt | { Star-forming galaxies have been predicted to contribute considerably to the diffuse gamma-ray background as they are guaranteed reservoirs of cosmic rays. Assuming that the hadronic interactions responsible for high-energy gamma rays also produce high-energy neutrinos and that $\mathcal{O}(100)$~PeV cosmic rays can be produced and confined in starburst galaxies, we here discuss the possibility that star-forming galaxies are also the main sources of the high-energy neutrinos observed by the IceCube experiment. First, we compute the diffuse gamma-ray background from star-forming galaxies, adopting the latest {\it Herschel} PEP/HerMES luminosity function and relying on the correlation between the gamma-ray and infrared luminosities reported by {\it Fermi} observations. Then we derive the expected intensity of the diffuse high-energy neutrinos from star-forming galaxies including normal and starburst galaxies. Our results indicate that starbursts, including those with active galactic nuclei and galaxy mergers, could be the main sources of the high-energy neutrinos observed by the IceCube experiment. We find that assuming a cosmic-ray spectral index of 2.1--2.2 for all starburst-like galaxies, our predictions can be consistent with both the {\it Fermi} and IceCube data, but larger indices readily fail to explain the observed diffuse neutrino flux. Taking the starburst high-energy spectral index as free parameter, and extrapolating from GeV to PeV energies, we find that the spectra harder than $E^{-2.15}$ are likely to be excluded by the IceCube data, which can be more constraining than the {\it Fermi} data for this population. } | \label{sec:introduction} %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% The existence of the extragalactic diffuse gamma-ray background (EGRB) has been revealed long ago~\cite{Fichtel1977}. Its importance is related to the fact that it encodes unique information about high-energy processes in the Universe. The EGRB has an isotropic sky distribution by its definition, and is contributed by gamma rays produced from leptonic processes or by cosmic rays interacting with the surrounding gas and radiation fields. It is most probably made of the superposition of contributions from unresolved extragalactic sources including star-forming galaxies, active galactic nuclei (AGN), gamma-ray bursts, intergalactic shocks produced by structure formation, etc.~(see Refs.~\cite{Dermer:2007fg,Inoue:2011bp,Collaboration:2010gqa,Ackermann:2012vca,Stecker:2010di,DiMauro:2013zfa}). Before {\it Fermi} and IceCube data being available, star-forming galaxies had been expected to contribute to the diffuse GeV-TeV gamma-ray background, see for example Refs.~\cite{sto+76,sj99,Pavlidou:2002va,tho+06}, and possibly to the cumulative sub-PeV neutrino background~\cite{lw06,Thompson:2006np,Lacki:2010vs,Murase:2013rfa,kat+13}. Their main emission component above GeV energies is due to cosmic-ray interactions with the diffuse gas. Diffuse emission from the Milky-Way galaxy (MW) is, in fact, the brightest feature of the gamma-ray sky. Gamma rays primarily arise from the decay of $\pi^0$ mesons produced in inelastic collisions of cosmic rays with the interstellar gas. As the MW, it is expected that other star-forming galaxies emit gamma rays through the same interactions providing substantial contribution to the EGRB. A special subset of star-forming galaxies is a less numerous, but individually more luminous population: The starburst galaxies. Starbursts undergo an epoch of star formation in a very localized region at an enhanced rate (i.e., $10$--$100\ M_\odot/$year) in comparison to normal galaxies such as the MW ($1$--$5\ M_\odot/$year). This activity is often triggered by galaxy mergers or by galactic bar instabilities where the dynamical equilibrium of the interstellar gas is disturbed and leads to the formation of high-density gas regions, usually at the center of the galaxy. For example, M82 and NGC 253 are located relatively nearby ($D \sim 2.5$--4.0~Mpc) and they can be defined as prototypical starbursts; each of them has an intense star-forming region in the center (with a radius of $200$~pc). These starbursts are expected to have high supernova rates of about 0.03--0.3~yr$^{-1}$ (see Ref.~\cite{Lacki:2010vs} for a detailed study on the characteristics of these sources). Observationally, EGRET detected only the MW and Large Magellanic Cloud~\cite{Abdo:2010pq}, but {\it Fermi} has also detected the Small Magellanic Cloud~\cite{Abdo:2010d}, the starburst galaxies M82 and NGC 253~\cite{Abdo:2010f} as well as starbursts with Seyferts NGC 4945 and NGC 1068~\cite{Abdo:2010g}, and some others more recently~\cite{Ackermann:2012vca,Hayashida:2013wha}. Our understanding of how galaxies form and evolve has increased dramatically over the last decade (see Ref.~\cite{Dunlop:2012pt} for a review). A wide range of surveys shows that the star formation rate density rises up to $z \sim 1$--2 and reveal that the bulk of the stellar mass density seen in the universe today was formed between $z \sim 1$ and $\sim$ 2. Understanding why the universe was much more active in the past and which processes or mechanisms drove galaxy evolution is one of the most important goals in the field of galaxy formation. In particular, one has to consider the energy absorbed by dust and re-emitted at longer wavelengths, in the infrared (IR) or sub-millimeter (sub-mm) range. Dust is responsible for obscuring the ultraviolet and optical light from galaxies. Since star formation is affected by dusty molecular clouds, far-IR and sub-mm data, where the absorbed radiation is re-emitted, are very important to provide a complete picture of the star formation history. Surveys of dust emission performed with former satellites exploring the Universe in the mid- and far-IR domain (IRAS, ISO, and \emph{Spitzer}) allowed the first studies of the IR-galaxy luminosity function (comoving number density of galaxies per unit luminosity range) up to $z \sim 2$~\cite{Rodighiero:2009up,Dole:2006,Hauser:2001xs,Puget:1996fx}. The detection of large numbers of high-$z$ sources, however, was not achievable before the \emph{Herschel Space Observatory}~\cite{Gruppioni:2013jna}, which has recently provided an estimate of the luminosity function up to $z\simeq4$. Strong evolution in both luminosity and density has been found, indicating that IR galaxies were more luminous and more numerous in the past~\cite{Gruppioni:2013jna,Devlin:2009qn,Bethermin:2012jd,Barger:2012st,Casey:2014hya}. {\it Herschel} even enabled, for the first time, the estimate of the luminosity functions of specific galaxy populations: normal galaxies, starbursts, and star-forming galaxies with obscured or low-luminosity AGNs, all of them contributing to the star-formation rate. We note, however, that our knowledge of the far-IR luminosity function is still affected by substantial uncertainties, because of source confusion and low detector sensitivity~\cite{Pilbratt:2010mv}. The IceCube neutrino observatory at the South Pole is presently the most sensitive instrument to uncover astrophysical neutrino sources in the TeV to PeV energy range. The IceCube Collaboration has recently reported evidence for extraterrestrial neutrinos, after the observation of three PeV neutrino cascades within three years of operation ~\cite{Aartsen:2013bka,IceCubetalk,Aartsen:2014gkd}. The recent data set consisting of $37$ events corresponds to a spectral excess with respect to the atmospheric background with a significance of more than $5 \sigma$~\cite{Aartsen:2014gkd}. This spectral excess is consistent with a diffuse neutrino flux with a $E^{-2}$ spectrum, possible break/cutoff at a few PeV, and isotropic sky distribution. The flux for each neutrino species is well described with the following fit based on the data between $30$~TeV and $2$~PeV: \begin{equation} \label{fitnu} E_{\nu}^2\ I_\nu(E_\nu) \simeq (0.95 \pm 0.3) \times 10^{-8}\ \mathrm{GeV}\ \mathrm{cm}^{-2}\ \mathrm{s}^{-1}\ \mathrm{sr}^{-1}\, \end{equation} within $1\sigma$ error bars. The energy distribution of the 37 neutrino events should provide additional clues about the candidate sources and several explanations have been indeed proposed to explain such high-energy events (see Refs.~\cite{wax13,Anchordoqui:2013dnh} for recent reviews). Standard fluxes of the prompt neutrino background from the decay of charmed mesons are too low to have a significant contribution at PeV energies~\cite{Enberg:2008te,Gaisser:2013ira}. A connection to cosmogenic neutrinos produced via the extragalactic background light seems unlikely~\cite{Laha:2013lka,Roulet:2012rv}, unless one assumes the optimistic extragalactic background disfavored by {\it Fermi} observations of gamma-ray bursts along with relatively low maximum proton energies~\cite{Kalashev:2013vba}. Early studies suggested a possible connection to the Glashow resonance~\cite{Bhattacharya:2011qu}, but this scenario was later disfavored by a follow-up analysis~\cite{Aartsen:2013bka}. More exotic models such as the PeV dark matter decay scenarios have been suggested too~\cite{Feldstein:2013kka,Esmaili:2013gha}. Galactic neutrino sources have been discussed, pointing out a possible association with the extended region around the Galactic center (see Refs.~\cite{Ahlers:2013xia,raz13} and references therein) or with the unidentified TeV gamma-ray sources~\cite{Fox:2013oza}. The Galactic halo has been discussed as one of the possible options~\cite{Ahlers:2013xia,Joshi:2013aua,Taylor:2014hya}. However, most of the observed events do not originate from the extended region around the Galactic center which makes the scenario of the Galactic origin slightly disfavored, and GeV-PeV gamma-ray observations have already put meaningful constraints on various Galactic possibilities including the Galactic halo scenario~\cite{Ahlers:2013xia,Murase:2013rfa}. The isotropic diffuse flux is most naturally explained with extragalactic sources. Among extragalactic scenarios, various PeV neutrino sources including gamma-ray bursts, peculiar supernovae, newborn pulsars, AGN, star-forming galaxies and intergalactic shocks had already been suggested before the discovery of the IceCube spectral excess. The observation can be most probably associated with various extragalactic $\sim100$~PeV cosmic-ray accelerators, e.g. low-power gamma-ray burst jets~\cite{Murase:2013ffa}, AGN~\cite{:2013fxa,Murase:2014foa}, star-forming galaxies including starbursts, galaxy mergers and AGN~\cite{lw06,Thompson:2006np,Lacki:2010vs, Murase:2013rfa,he+13,liu+14,kat+13,Kashiyama:2014rza,Anchordoqui:2014yva,Chang:2014hua}, intergalactic shocks and active galaxies embedded in structured regions~\cite{Murase:2013rfa}. Mixed scenarios of Galactic and extragalactic neutrino sources have also been discussed~\cite{Ahlers:2013xia,raz13,Fox:2013oza,Joshi:2013aua,Murase:2014foa,Padovani:2014bha}. Since the star-forming galaxies are considered to be one of the main source classes of the diffuse gamma-ray background, we here consider them as a potential source population of the high-energy neutrinos observed by IceCube and compare our estimations with the observed IceCube flux. Making such a multi-messenger connection is one of the keys for revealing the origin of the diffuse neutrino flux observed by IceCube. As shown in Ref.~\cite{Murase:2013rfa}, the multi-messenger connection based on the measured neutrino and gamma-ray data can provide crucial ways to constrain various extragalactic scenarios such as star-forming galaxies. We rely our computations on the latest IR luminosity function provided by {\it Herschel} PEP/HerMES up to $z \simeq 4$~\cite{Gruppioni:2013jna} as well as the well calibrated scaling relation between gamma-ray luminosity and IR luminosity~\cite{Ackermann:2012vca}. Compared with previous studies on the EGRB estimates due to star-forming galaxies~\cite{Pavlidou:2002va,tho+06,Fields:2010bw,Makiya:2010zt,Lacki:2012si}, we adopt the empirical relation obtained in the {\it Fermi} era as well as the luminosity functions of different populations of galaxies (normal galaxies, starbursts, galaxies containing obscured/low-luminosity AGN), separately, which became possible thanks to {\it Herschel}'s great statistics of various populations covering wide range of the IR spectrum. For our \emph{canonical} model, where we assume that the spectral index of both gamma rays and neutrinos is $2.2$ from starburst-like sources, we show that a significant fraction ($\sim$20--50\%) of the IceCube flux may be explained with this source population given that the extrapolation from GeV to PeV energies is realized. If the starburst population features harder spectra than $E^{-2.15}$, then the IceCube data yield more stringent constraints on such a scenario than the {\it Fermi} data, allowing us to exclude these models. This paper is organised as follows. We first calculate the EGRB intensity based on the {\it Herschel} PEP/HerMES luminosity function, and compare the EGRB estimation with the {\it Fermi} data in Sec.~\ref{sec:IRluminosityfunction}. We then present the associated diffuse neutrino background in comparison with the IceCube data in Sec.~\ref{sec:gammanu}. We constrain the injection spectral index of starburst-like galaxies as well as their abundance in Sec.~\ref{sec:constraints}. In Sec.~\ref{sec:issues}, we discuss caveats and issues of the star-forming galaxy scenario for the observed high-energy neutrinos, which should be addressed in the future. We finally give summary and discuss perspectives in Sec.~\ref{sec:conclusions}. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% | \label{sec:conclusions} %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% The importance of the EGRB is related to the fact that it encodes unique information about the high-energy processes in the universe, it has an isotropic sky distribution and is dominated by gamma rays induced by cosmic-ray protons and electrons interacting with the target gas and radiation fields. Although main contributions of the diffuse EGRB are likely to come from star-forming galaxies and blazars, such a diffuse background should also include other sub-leading contributions from unresolved extragalactic sources in the different energy range, such as radio galaxies, gamma-ray bursts, intergalactic shocks produced by structure formation, and dark-matter annihilation. Star-forming galaxies are expected to largely contribute to the $0.3$--$100$~GeV gamma-ray background, because of interactions of cosmic-rays with the diffuse gas. Besides normal galaxies, a special subset of star-forming galaxies are the less numerous, but individually more luminous, starburst galaxies. The starburst galaxies undergo an epoch of star formation in a very localized region at an enhanced rate in comparison to the normal galaxies, due to galaxy mergers or bar instabilities. According to the classification done by {\it Herschel} in terms of the luminosity functions, the largest fraction of star-forming galaxies is made by galaxies containing low luminosity or dim AGN, and such family could have an energy spectrum similar to normal galaxy or to starburst galaxies according to the cases. Our understanding of galaxy formation has dramatically improved in the last few years. In particular, the {\it Herschel} PEP/HerMES survey has recently provided an estimate of the IR luminosity function up to $z \simeq 4$, finally allowing to explore the galaxy composition for $z \ge 2$, suggesting that galaxies containing AGN are largely abundant and that the starburst fraction can reach up to $30\%$ of the total at high $z$. We have provided an estimation of the EGRB adopting IR data at $z \ge 2$ through the luminosity function provided in~Ref.~\cite{Gruppioni:2013jna}. The luminosity function indeed provides a powerful tool to probe the distribution of galaxies over cosmological time, since it allows to access the statistical nature of galaxy formation and their evolution. Our estimation is compatible with the {\it Fermi} data in the 0.3--30~GeV range including the astrophysical uncertainties. In our canonical model, the sum of NG and SB components leads to $\Gamma_{\rm SF}\sim2.4$, where we find that high-energy data are not explained due to the EBL attenuation. We identify the very-high-energy gamma-ray excess, which may indicate another population like BL Lac objects, although quantitative details are affected by EBL uncertainties. High-energy neutrinos are also expected to be emitted by these sources together with gamma rays. Assuming that at least $\sim100$~PeV cosmic rays are accelerated and confined in starburst galaxies, we estimated the expected diffuse neutrino background from star-forming galaxies, and find that it can be in agreement with the IceCube data, especially at low energies. Although such estimation needs to be taken with caution since better statistics of the IceCube high-energy events is required, the disagreement at high energies might suggest that the observed IceCube flux is given by the contribution of other sources such as gamma-ray bursts and AGN. In particular, in our star-forming galaxy scenario, we found that for $\Gamma_{\mathrm{SB}} \simeq 2.15$ both {\it Fermi} and IceCube data are consistently explained. The {\it Fermi} and IceCube data could constrain the abundance and the injection spectral index of the fraction of galaxies containing AGN that behave similarly to starbursts. We find that hard spectra for SB and SB-like SF-AGN (i.e., $\Gamma_{\rm SB}\lesssim2.1$) are excluded from present IceCube data and, as expected, the data allow more abundant SB-like galaxies for softer spectra, since they give a lower contribution to the high energy tail of the spectrum. In conclusion, our understanding of how galaxies evolve has dramatically increased over the last decade and, after the new discovered IceCube high-energy neutrinos, we have shown as a powerful multi-messenger connection between gamma rays and neutrinos can provide important new ways to constrain star-forming galaxy distribution as well as their spectral parameters. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% | 14 | 4 | 1404.1189 |
1404 | 1404.0213_arXiv.txt | We present the results of a search for companions to young brown dwarfs in the Taurus and Chamaeleon~I star forming regions ($\tau\sim1$ and 2--3~Myr). We have used the Wide Field Planetary Camera 2 on board the {\it Hubble Space Telescope} to obtain F791W and F850LP images of 47 members of these regions that have spectral types of M6--L0 ($M\sim0.01$--0.1~$M_\odot$). An additional late-type member of Taurus, FU~Tau (M7.25+M9.25), was also observed with adaptive optics at Keck Observatory. We have applied PSF subtraction to the primaries and have searched the resulting images for objects that have colors and magnitudes that are indicative of young low-mass objects. Through this process, we have identified promising candidate companions to 2MASS J04414489+2301513 ($\rho=0\farcs105$/15~AU), 2MASS J04221332+1934392 ($\rho=0\farcs05$/7~AU), and ISO~217 ($\rho=0\farcs03$/5~AU). We reported the discovery of the first candidate in a previous study, showing that it has a similar proper motion as the primary through a comparison of astrometry measured with WFPC2 and Gemini adaptive optics. We have collected an additional epoch of data with Gemini that further supports that result. By combining our survey with previous high-resolution imaging in Taurus, Chamaeleon~I, and Upper Sco ($\tau\sim10$~Myr), we measure binary fractions of 14/93 = $0.15^{+0.05}_{-0.03}$ for M4--M6 ($M\sim0.1$--0.3~$M_\odot$) and 4/108 = $0.04^{+0.03}_{-0.01}$ for $>$M6 ($M\lesssim0.1$~$M_\odot$) at separations of $>10$~AU. Given the youth and low density of these three regions, the lower binary fraction at later types is probably primordial rather than due to dynamical interactions among association members. The widest low-mass binaries ($>100$~AU) also appear to be more common in Taurus and Chamaeleon~I than in the field, which suggests that the widest low-mass binaries are disrupted by dynamical interactions at $>$10~Myr, or that field brown dwarfs have been born predominantly in denser clusters where wide systems are disrupted or inhibited from forming. | As with stars at higher masses, the binary properties of low-mass stars and brown dwarfs may provide insight into their formation and dynamical evolution \citep{duc13}. Multiplicity at low masses has been characterized primarily through high-resolution imaging in the solar neighborhood \citep{koe99,mar99,rei01,bou03,bur03,clo03} and the nearby young clusters and associations \citep{neu02,mar03,luh05ic,kra05,kra06,kon07,bil11,kra12}. These surveys have found that the binary fractions and the separations of binaries decrease and the mass ratios increase from stars to brown dwarfs \citep{bur07ppv,kra12}. Although most binary brown dwarfs have small separations ($a<20$~AU), a few wide systems have been uncovered \citep{luh04bin,cha04}. The low binding energies of these wide binaries would tend to suggest that dynamical interactions did not play a role in their formation \citep{rc01}, although a few ejected brown dwarfs may be captured into wide systems in denser clusters \citep{bat05}. The dependence of the binary properties of low-mass stars and brown dwarfs on age and star-forming environment is not well-constrained by existing data. As a result, it is unclear whether wide low-mass binaries are frequently disrupted by interactions with stars, either in their natal clusters or in the Galactic field, and how the initial conditions of star formation influence binarity at low masses. The Taurus and Chamaeleon~I star-forming regions are promising sites for providing some of the data needed to investigate these issues. They are among the nearest star-forming regions (140 and 165~pc), young enough that dynamical interactions are minimized (1 and 2--3~Myr), most of their members have relatively low extinction ($A_V\lesssim4$), and they have been searched thoroughly for substellar members \citep{ken08,luh08cha}. These two regions also offer the opportunity for characterizing the multiplicity produced by low-density star-forming conditions, which can be compared to measurements in richer and denser clusters at the same age as well as older populations in open clusters and the solar neighborhood. High-resolution imaging has been previously applied to low-mass members of Taurus and Chamaeleon~I \citep{neu02,kra06,ahm07,kon07,luh07cha,laf08,kra12}, which has included 45 primaries with spectral types later than M6 ($M\lesssim0.1$~$M_\odot$). To improve upon the statistical accuracy of those multiplicity measurements, we have performed an imaging survey that includes most of the remaining known late-type members of Taurus and Chamaeleon~I using the {\it Hubble Space Telescope} ({\it HST}) and Keck Observatory, resulting in a sample of 73 primaries in these regions later than M6 for which high-resolution data are now available. In this paper, we describe the sample selection and observing strategy in our survey (Section \ref{sec:obs}) and our analysis of the resulting images (Section \ref{sec:images}). We then identify the most promising candidate companions in our data and combine our sample with those of previous surveys to measure the binary fraction as a function of spectral type (Section~\ref{sec:cand}). We conclude by discussing the implications of our survey for measurements of multiplicity at low masses and for the formation of brown dwarfs (Section~\ref{sec:disc}). | \label{sec:disc} Taurus, Chamaeleon~I, and Upper Sco contain the largest samples of young low-mass stars and brown dwarfs that have been imaged at high resolution, providing the best available statistical constraints on low-mass multiplicity at ages of $\lesssim10$~Myr. We first examine the dependence of the binary fractions in Table~\ref{tab:fraction} on spectral type of the primary. For each of the three regions, the wide binary fraction ($>$10~AU) is significantly lower at $>$M6 than at M4--M6. Trends of this kind have been detected previously in subsets of the data we have compiled \citep{kra05,kra06,bil11,kra12}, in samples of members of Taurus, Chamaeleon~I, and Upper Sco at higher masses \citep{laf08,kra09}, and in the solar neighborhood \citep[][references therein]{bur07ppv}. Given the youth and low density of the regions in question, particularly Taurus, this dependence on spectral type is very likely primordial rather than due to dynamical interactions among members of each region. We also can examine the data in Table~\ref{tab:fraction} for differences among Taurus, Chamaeleon~I, and Upper Sco. For each range of spectral types, the binary fractions do not show any statistically significant differences between the regions. Comparing these binary fractions to data for field stars and brown dwarfs is more problematic because it is difficult to ensure that young and old samples encompass the same ranges of primary masses, and because a given young cluster may not represent the predominant star-forming environment for the field. Nevertheless, it is useful to compare the frequency of the widest brown dwarf binaries. Taurus and Chamaeleon~I each contain one known binary brown dwarf with a separation greater than 100~AU \citep{luh04bin,luh09fu}. Although the Upper Sco sample that we have defined for Table~\ref{tab:fraction} does not have any pairs wider than 100~AU, a few examples have been found among other brown dwarfs in the association \citep{all06,jay06,clo07,luh07oph,bej08}. Thus, the data for Taurus, Chamaeleon, and Upper Sco indicate a binary fraction of a few percent for these wide binary brown dwarfs. In comparison, only one pair of this kind has been found among the several hundred known late-L and T dwarfs in the field \citep{burn10,sch10}. This implies that dynamical interactions with cluster members or field stars at $>$10~Myr disrupt the widest binary brown dwarfs, or that most field brown dwarfs are born under conditions different from those in Taurus, Chamaeleon~I, and Upper Sco, perhaps in denser clusters where very wide binaries are disrupted or prevented from forming. | 14 | 4 | 1404.0213 |
1404 | 1404.2811_arXiv.txt | CCD multi-band light curves of the neglected eclipsing binaries V405~Cep, V948~Her, KR~Mon and UZ~Sge were obtained and analysed using the Wilson-Deninney code. New geometric and absolute parameters were derived and used to determine their current evolutionary state. V405~Cep, V948~Her and KR~Mon are detached systems with their components in almost the same evolutionary stage. UZ~Sge is a classical Algol system with a tertiary companion around it. Moreover, since the systems V405~Cep, V948~Her and UZ~Sge contain an early type component, their light curves were examined for possible pulsation behaviour. | \label{1} Eclipsing binary systems offer unique conditions to measure fundamental parameters of stars, such as stellar masses, radii and luminosities, which are of great importance in studies of stellar structure and evolution. The main purposes of this work are: a) the derivation of the geometric and photometric parameters of each system through light curve (hereafter LC) analysis, b) the approximate calculation of the absolute parameters of their components and c) a description of their current evolutionary status. In addition, frequency analysis search was performed for possible pulsational behaviour in the candidate systems for including a $\delta$~Sct component. Concluding, in the present study we selected the following three, previously neglected, systems which are lacking accurate photometric measurements and/or modern LC analysis in order to derive for the first time their main physical characteristics. \textbf{V405~Cep:} This system ($B$=8.95~mag, $P$=1.37374$^{\rm d}$) was discovered as a variable by $Hipparcos$ mission \citep{ES97} and has an Algol-type LC. The only available information for this binary concerns its A2 spectral type \citep[cf.][]{WR03,KH01}, while \citet{SO06} included it in their list of candidate systems for including pulsating components. \textbf{V948~Her:} The system's Algol-type variability was discovered by $Hipparcos$ mission \citep{ES97}. It has a period of 1.27521$^{\rm d}$, a $B$-magnitude of 9.26 and its spectral type is listed as F2 in many catalogues \citep[cf.][]{WR03,BU04}. The system is also candidate for containing a $\delta$~Sct component \citep{SO06}. \textbf{KR~Mon:} This binary was discovered as a variable by \citet{ZI52}. Its period is 1.15097$^{\rm d}$ and its brightness has a magnitude of 11.7 in $B$-filter. The \textit{GCVS} catalogue \citep{SA11} and the \textit{MK Catalogue of Stellar Spectral Classifications} \citep{SK10} list it as a G3V type star. LCs of the system in $V$ and $I$ filters were obtained by the $ASAS$ project \citep{PO05}, but no analysis has been made so far. \textbf{UZ~Sge:} The binary nature of this Algol-type system ($B$=11.4~mag, $P$=2.21574$^{\rm d}$) was reported by \citet{GS39}. Its spectral type is given either as A3V \citep{HA84,SK90} or A0V \citep{BD80,BU84}. \citet{BD80} and \citet{SK90} calculated the absolute parameters of its components but they resulted in different values. There are many minimum timings available in the literature since the discovery of the system. \citet{LN08} and \citet{ZA08} analysed its orbital period changes and concluded that, a third body with minimal mass of $\sim$0.7~M$_\odot$ may exist in a wide orbit. \citet{BU09} obtained and analysed $V$ and $R$ LCs of the system and resulted in two possible geometrical configurations with different mass ratios. | \label{6} Complete LCs of the eclipsing systems V405~Cep, V948~Her, KR~Mon and UZ~Sge were obtained and analysed using well known modern techniques. The results were used to calculate the absolute parameters of their components and make an estimate of their present evolutionary status. V405~Cep is a detached system with its primary component almost in the middle of ZAMS-TAMS limits, while its secondary was found to be closer to ZAMS. According to the mass values of these stars, it seems that the more massive component is more evolved that the less massive one. This situation is in agreement with the stellar evolutionary models. However, we notice that the stars show large radius difference, so, it is expected that the primary will fill its Roche lobe first as it evolves faster and, therefore, it will become a mass looser. For V948~Her it was found that its components are more evolved, reaching the TAMS, but they are well inside their Roche lobes. It is noticed that, although its components have a large difference in their masses, they are almost in the same evolutionary status. This fact comes in disagreement with our knowledge for stellar evolution, assuming a simultaneous birth of the components. This discrepancy could be explained with the scenario of past mass transfer from the present secondary to the primary, or less likely with a rapid mass loss (e.g. due to very strong stellar winds) of the secondary. On the other hand, the picture of KR~Mon is different. Its components have almost equal temperatures, masses and sizes, but none of them has reached its critical Roche volume so far. Both of them are slightly evolved and lie outside the TAMS limit. However, they are very close in filling their lobes, therefore they will probably come into contact in the future. Of course, the latter needs further justification with future observations, e.g. a period study showing if they are approaching each other. UZ~Sge was found to be a classical Algol with the less massive and cooler component filling its Roche lobe. Its primary is close to ZAMS and was revealed as a $\delta$~Sct type pulsator, whose pulsating phase must have started relatively recently, if we consider both its dominant frequency value (i.e. fast pulsator) and its present evolutionary status \citep{LI12}. Therefore, according to the definition given by \citet{MK04} it can also be considered as oEA star (i.e. a classical Algol system including an oscillating MS component of A-F spectral type). The secondary component of the system is far beyond the MS edge and it must have been transferring a great part of its mass to the primary according to its present mass and evolutionary stage. Therefore, in the past it was very probably the more massive star of the system, evolved faster, reached its critical Roche limit and after that it started losing mass up to the present. However, according to the period analysis of the system \citep{LN08,ZA08} no secular period change that can be connected with the mass transfer process \citep{HI01} was found. Probably the system is at slow mass-accretion stage \citep{MK03} with a rate that cannot be detected with the current time coverage of minima timings. \citet{LN08} and \citet{ZA08} revealed also the possibility of a tertiary component's existence with a minimal mass of $\sim$0.7~M$_\odot$ and a mass function $f(m_3)=0.031$~M$_\odot$. From the present LC analysis we found a third light contribution of $L_{3,{\rm LC}}\sim$7.5\% to the total luminosity of the system. Assuming the MS nature of the third component, its luminosity based on the Mass-Luminosity relation for dwarf stars ($L\sim M^{3.5}$) can be derived. Given the absolute luminosities of the binary's members (see Table \ref{tab3}), we can estimate the expected light contribution ($L_{3,{\rm O-C}}$) of the potential third star by using the following formulae: \begin{equation} L_{3,{\rm O-C}} (\%)=100 \frac{M_{3,{\rm min}}^{3.5}}{L_1+L_2+M_{3,{\rm min}}^{3.5}} \end{equation} According to this hypothesis, we found that the third body's luminosity percentage is 1.4\%, that means less than the observed one. This difference, although it is large, it is based both on the minimal mass of the third body (coplanar orbit) and on its MS nature. Hence, if we use the mass function equation of the third body \citep[cf.][]{TO10}: \begin{equation} f(M_3) = \frac{1}{P_3^2}\left[\frac{173.145 A}{\sqrt{1 - e_3^2 \cos^2 \omega}} \right]^3 = \frac{(M_3 \sin i_3)^3}{(M_1+M_2+ M_3)^2} \end{equation} with the wide orbit's period $P_3$ in yr, and the LITE amplitude $A$ in days, we can see that, with $i_3\sim40^\circ$, we get $M_3\sim1.2$~M$_\odot$, a value satisfying the observed luminosity contribution. As an alternative, we can also assume that the third body is an evolved star providing more light than a MS star with the same mass. To sum up, we conclude that indeed there must be another component around the system, but its mass and nature remain open questions. The mass ratio values for V405~Cep and V948~Her show deviation from the average of the `well known' detached binaries from the list of \citet{IB06}. However, that list contains only 74 cases of well studied detached binaries, therefore many deviations are expected to be found. Similar systems with relatively small mass ratio values are e.g. AL~Gem and UU~Leo \citep{LI11a}, EL~Vel \citep{ZA11}, GSC~4589-2999 \citep{LI11b} and AR~Lac \citep{IB06}. Moreover, the LCs of V405~Cep and V948~Her present total eclipses (both ones for V405~Cep and the secondary one for V948~Her), hence, the mass ratios can be considered of well determined using the `$q$-search' method \citep{TW05}. However, the error values of $q$ for these systems are not that realistic, given that their `$q$-search' diagrams show a shallow minimum (see Fig.~1), therefore, they can be considered only as formal ones. The parameters of KR~Mon and UZ~Sge are well inside the limits of detached and semi-detached systems, respectively \citep{IB06}. Radial velocity measurements are certainly needed for deriving more accurate the absolute parameters of these close binaries and, therefore, obtaining a more realistic view of them. | 14 | 4 | 1404.2811 |
1404 | 1404.7575_arXiv.txt | A maximum jet efficiency line $R\sim 25$ ($R=L_{\rm jet}/L_{\rm bol}$), found in FRII radio galaxies by Fernandes et al., was extended to cover the full range of jet power by Punsly. Recent general relativistic magnetohydrodynamic (GRMHD) simulations on jet formation mainly focused on the enhancement of jet power. In this work, we suggest that the jet efficiency could be very high even for conventional jet power if the radiative efficiency of disk were much smaller. We adopt the model of a thin disk with magnetically driven winds to investigate the observational high efficiency jets in FRII radio galaxies. It is found that the structure of a thin disk can be significantly altered by the feedback of winds. The temperature of disk gradually decreases with increasing magnetic field; the disk density, surface density and pressure also change enormously. The lower temperature and higher surface density in inner disk result in the rapid decrease of radiative efficiency. Thus, the jet efficiency is greatly improved even the jet power is conventional. Our results can explain the observations quite well. A theoretical maximum jet efficiency $R \sim 1000$ suggested by our calculations is large enough to explain all the high jet efficiency in observations even considering the episodic activity of jets. | Relativistic jets are common characters of active galactic nuclei (AGNs). According to the different morphology of their radio structures, the radio galaxies can be divided into two classes, i.e., FRI (defined by edge-darkened radio lobes) and FRII (defined by edge-brightened radio lobes and hot spots) \citep{f1974}. Several recent observations discovered that the jet efficiency (defined as $R=L_{\rm jet}/L_{\rm bol}$) of some luminous lobe-dominated FRII could be very high, where $L_{\rm jet}$ and $L_{\rm bol}$ are the jet power and the bolometric luminosity of AGNs, respectively \citep*[][hereafer P11]{m2011,f2011,p2011}. A maximum jet efficiency $R \sim 25$ was found by \citet{f2011}, where they adopted a complete sample of the most powerful radio galaxies at redshift $z\sim1$. P11 extended the results of \citet{f2011} and found that the maximum jet efficiency line can cover the full four order of magnitudes of jet power by including other samples, such as, a blazar sample in \citet{g2010}, a small sample of X-ray cavities in \citet{m2011}, etc. There are mainly two most popular jet formation mechanisms so far, i.e., the Blandford-Znajek (BZ) process \citep{b1977} and the Blandford-Payne (BP) process \citep{b1982}. In BZ process, the rotating energy of a black hole can be extracted to power a jet by the large scale magnetic fields maintained by an accretion disk. But in BP process, the jet power comes from the rotating disk itself instead. The bolometric luminosity of a radio galaxy can be expressed as $L_{\rm bol}=\eta_{\rm th} \dot{M}c^2$, where $\eta_{\rm th}$ is the radiative efficiency of the accretion disk. If we specify the jet power as a function of accretion rate, $L_{\rm jet}=\eta_{\rm Q} \dot{M}c^2$ (where $\eta_{\rm Q}$ is the jet production efficiency), the jet efficiency $R(=L_{\rm jet}/L_{\rm bol}=\eta_{\rm Q}/\eta_{\rm th})$ will be decided by both $\eta_{\rm Q}$ and $\eta_{\rm th}$. The obviously different accretion rate between FRI and FRII radio galaxies implies that they should have different accretion model \citep{l1996,g2001,x2009}. In the general picture of FRII radio galaxies, jet is supposed to be launched from a ridiatively efficient accretion disk , where $\eta_{\rm th}$ varies from about $0.06$ to $0.4$ for a non-rotating black hole and a extreme Kerr black hole, respectively. Recent GRMHD simulations on jet formation mainly focused on the improvement of jet power $L_{\rm jet}$ in order to explain the observed high jet production efficiency $\eta_{\rm Q}$. The jet production efficiency for a magnetically-arrested-disk (MAD) can reach $\sim 30\%$ and $140\%$ for $a=0.5$ and $0.99$, respectively \citep{t2011}, which implies that the black hole spin parameter $a$ may play a key role in the formation of jets \citep{m2005,t2010,t2011}. However, the observed $\eta_{\rm Q}$ is based on an assumption that the radiative efficiency $\eta_{\rm th}\sim 0.1$, which is suggested from observational constraints on the growth of massive black holes \citep*[e.g.,][]{y2002}. Thus, if the radiative efficiency of a disk were much smaller than $0.1$, it is possible to get a high jet efficiency $R$ even for conventional jet production efficiency $\eta_{\rm Q}$. \citet{l2012} and \citet{l2014} do seem to provide a possible way for this picture. They produced a model for a thin disk with magnetically driven winds/jets, in which the angular momentum and energy carried away by jets are properly included. It was found that the disk properties can be changed significantly by the feedback of jets. For example, the temperature of a thin disk with jets is obviously lower compared with that of a standard thin disk \citep*[see figure 1,][]{l2012} for the reason that a large fraction of gravitational energy released in the disk is carried away by jets. This will directly result in the decrease of bolometric luminosity (and $\eta_{\rm th}$) of the accretion disk and the increase of jet efficiency $R$. In this work, we will detailedly investigate the structure of a thin disk with winds and the high efficiency jets in powerful lobe-dominated FRII radio galaxies. | \label{sec:summary} The basic physics of the jet formation models adopted in this work are the same as the previous works of \citet{g1997} and \citet{l1999}. However, their estimates of the magnetic field strength of a disk are based on the conventional accretion disk models without magnetic field, which is a good approximation in the weak magnetic field case as the disk structure has not been altered significantly by the field. But that assumption becomes invalid if the field is strong. The high power carried by the jets has greatly changed the structure of a thin disk (figure \ref{properties}, \ref{pressure}). Except for the angular velocity, which is close to Keplerian angular velocity for a thin disk \citep{l2014}, all other disk properties (temperature, density, surface density, pressure) change a lot compared with the standard disk. The magnetic torque $T_{\rm m}$ is found to dominate over the viscous torque $T_{\rm vis}$ (see figure \ref{properties}d). The temperature decreases and the density increases significantly with decreasing $\beta_{\rm p}$. It is the increase of surface density and the decrease of disk temperature in the inner disk region that result in the very low radiative efficiency in the disk. The obviously lower temperature compared with that of a standard thin disk even leads to the disappearance of the inner disk dominated by radiative pressure (figure \ref{pressure}). Thus, the thin disk becomes both thermally and viscously stable on the presence of disk winds \citep{l2014}. Previous efforts mainly focused on the improvement of jet power $L_{\rm jet}$. But indeed, if the accumulation of magnetic flux in the inner region of accretion disk is considered \citep{n2003,porth2010,porth2011,t2011,m2012,s2013}, not only can the jet production efficiency be improved significantly, the radiative efficiency can also decrease a lot. Thus, a very high jet production efficiency $\eta_{\rm Q}$ isn't always needed. The jet efficiency could be quite 'normal' if the radiative efficiency decreases significantly (see figure \ref{radiative}, \ref{jet}). According to our model, the reason for the high jet efficiency is that most of released gravitational energy in the disk is carried away by jets, which results in the very low radiative efficiency. A theoretical maximum jet efficiency $R\sim 1000$ is found, which is large enough even we take the episodic activity of jets into account. Compared with the observational results in P11 (the dashed line in figure \ref{ljet}), our study indicates that $10< \beta_{\rm p}<20$ is required for $\alpha=0.1$ and $B_{\phi}=0.1 B_{\rm p}$. But if we consider smaller $\alpha$ and larger proportion of $B_{\phi}$ in the field, the magnetic torque could be more dominant \citep{l2014}. Thus the required field could be much weaker (larger $\beta_{\rm p}$). From figure \ref{ljettheta}, it seems that the jet efficiency has a positive relation with $\theta$. But if we consider the evolution of field in a thin disk, the smaller $\theta$ may represent a stronger magnetic field \citep{l1994a,c2013}, which should thus help to improve the jet efficiency. The large scale magnetic field plays a key role in the formation of jet. But how it forms is still an unsolved problem. A popular mechanism is that the field lines can be dragged inwards with the accretion of gas. The magnification of field seems to be hard in a thin disk because the speed of turbulent diffusion is faster than that of advection \citep{v1989,l1994a}. But when taking the magnetically driven winds into account \citep{c2013,l2014}, the field may be effectively magnified even for a thin disk with very weak initial field ($\beta_{\rm p}\sim 10^3$). However, a balance of magnetic field advection and diffusion is required in order to avoid the formation of a MAD. How such a balance can be achieved and kept stable are still unclear \citep*[e.g.,][]{c2002b,c2013,b2012}. \citet{l1994b} argued that a disk-wind system may be unstable if its angular momentum is taken away by magnetic torque only. The reason is that if there is a perturbation which increases the radial velocity, the inclination angle of the field will become smaller, which in turn increases the mass-loss rate and results in a higher radial velocity. Nevertheless, the linear stability analysis given by \citet{c2002b} suggested that a disk could be stable if the field is weak enough. \citet{l1994b} also stated that they didn't do a global calculation on the disk. Thus, a global time-dependent study should be necessary in order to investigate this problem, which is beyond the scope of this work. The formation of large scale magnetic field also depends on the initial strength and morphology of magnetic field as indicated by some MHD simulations \citep*[e.g.,][]{i2003,b2008,m2012}, which is still an open issue at present. | 14 | 4 | 1404.7575 |
1404 | 1404.1638.txt | We present polarisation properties at $1.4\,$GHz of two separate extragalactic source populations: passive quiescent galaxies and luminous quasar-like galaxies. We use data from the {\it Wide-Field Infrared Survey Explorer} data to determine the host galaxy population of the polarised extragalactic radio sources. The quiescent galaxies have higher percentage polarisation, smaller radio linear size, and $1.4\,$GHz luminosity of $6\times10^{21}<L_{\rm 1.4}<7\times10^{25}\,$W Hz$^{-1}$, while the quasar-like galaxies have smaller percentage polarisation, larger radio linear size at radio wavelengths, and a $1.4\,$GHz luminosity of $9\times10^{23}<L_{\rm 1.4}<7\times10^{28}\,$W Hz$^{-1}$, suggesting that the environment of the quasar-like galaxies is responsible for the lower percentage polarisation. Our results confirm previous studies that found an inverse correlation between percentage polarisation and total flux density at $1.4\,$GHz. We suggest that the population change between the polarised extragalactic radio sources is the origin of this inverse correlation and suggest a cosmic evolution of the space density of quiescent galaxies. Finally, we find that the extragalactic contributions to the rotation measures (RMs) of the nearby passive galaxies and the distant quasar-like galaxies are different. After accounting for the RM contributions by cosmological large-scale structure and intervening Mg\,{II} absorbers we show that the distribution of intrinsic RMs of the distant quasar-like sources is at most four times as wide as the RM distribution of the nearby quiescent galaxies, if the distribution of intrinsic RMs of the WISE-Star sources itself is at least several rad m$^{-2}$ wide. | \citet{Tucci2012} recently found that the intrinsic percentage polarisation of extragalactic radio sources (ERS) at frequencies $\ge 20\,$GHz is between $2-5\,$per cent, independent of flux density. These results were confirmed by \citet{Massardi2013} using the Australia Telescope $20\,$GHz (AT20G) Survey, while \citet{Sadler2006} suggest that there is a trend that fainter sources tend to have higher percentage polarisation. This anti-correlation between percentage polarisation and total flux density has also been suggested at $1.4\,$GHz by \citet{Mesa2002}, \citet{Tucci2004}, \citet{Taylor2007}, \citet{Subrahmanyan2010}, and \citet{Grant2010}. Recently, \citet{Hales2013} found no evidence for this trend and attribute the previous results to selection effects consistent with the reasoning by \citet{Massardi2013}. As a result the anti-correlation of percentage polarisation with total flux density, if it exists, remains a mystery. The first studies of increasing percentage polarisation with decreasing flux density came from \citet{Mesa2002} and \citet{Tucci2004} who both suggested a population change of ERS at fainter flux densities was the cause. \citet{Taylor2007} went on to suggest that the cause was a result of a change in the fraction of radio quiet active galactic nuclei (AGN). Most recently, \citet{Rudnick2014} examined the polarisation properties of radio sources down to $S_{\rm 1.4\,GHz} > 15\,\mu$Jy in the GOODS-N field and suggest a population change around a polarised flux density of $1\,$mJy. Studies into the intrinsic properties of polarised ERS by \citet{Banfield2011} show a trend of increasing percentage polarisation with decreasing luminosity and no trend with redshift, later confirmed by \citet{Hammond2012}. \citet{Subrahmanyan2010} suggest that this anti-correlation between percentage polarisation and total flux density is likely to be a transition from FRII-dominated to FRI-dominated populations, while the results by \citet{Grant2010} imply that the higher percentage polarisation may be originating in the lobe-dominated structure and not in beamed BL Lac objects. However, \citet{Shi2010} found no dependence on ERS environment when comparing highly polarised ($> 30\,$per cent) ERS with their low polarised counterparts. \citet{Shi2010} went on to suggest that intrinsic properties of magnetic field ordering, thermal plasma density, and magnetic field orientation to the line of sight are the root cause for highly polarised ERS. In this paper we present an analysis of 1.4~GHz polarised ERS in combination with optical spectroscopic data in order to explore this anti-correlation between percentage polarisation and flux density. We probe polarised radio emission out to high-redshifts and examine the magnetic fields within different ERS populations. We outline the sample selection in Section \ref{sec:sample} and the nature of the polarised sources is discussed in Section \ref{sec:nature}. Section \ref{sec:seleff} describes the selection effects of our data, we discuss our findings in Section \ref{sec:dis}, and conclusions are presented in Section \ref{sec:conc}. The cosmological parameters used throughout this paper are: $\Omega_{\rm \lambda} = 0.7$; $\Omega_{\rm M} = 0.3$; and $H_{0} = 70$~kms$^{-1}$Mpc$^{-1}$. We define the spectral index $\alpha$ as $S\propto \nu^{\alpha}$. %------------------------------------------------------------------------- % MAGNETIC FIELDS %------------------------------------------------------------------------- | \label{sec:conc} Using the \citet{Hammond2012} catalogue of Faraday rotation measures and redshifts for 4003 ERS detected at $1.4\,$GHz, we have shown that polarised radio sources split into two types of host galaxies at two separate redshift ranges, as such the two populations are investigated separately and a larger sample of polarised radio sources is required to examine if one population evolves into the other population. We find the following: \vskip 0.2cm \noindent (1) the anti-correlation between percentage polarisation and total flux density is real as the percentage polarisation depends on WISE mid-infrared colour; \vskip 0.2cm \noindent (2) the polarised ERS separate clearly into two infrared-selected objects: WISE--Star sources that are low-redshift, low-radio-luminosity elliptical galaxies, and WISE--AGN which are high-redshift, high-radio-luminosity quasar-like galaxies; \vskip 0.2cm \noindent (3) our sample has a larger number of quiescent galaxies than \citet{Hales2013}, suggesting that the inconsistency between the data sets is an indication of cosmic evolution of the space density of quiescent galaxies; \vskip 0.2cm \noindent (4) we suggest that the difference in the percentage polarisation of radio galaxies originates from the environment of the host galaxy. Our WISE--AGN population is consistent with HzRGs in denser environments where depolarisation is more severe compared to the WISE--Star sources that are not very active; \vskip 0.2cm \noindent (5) we find that the extragalactic RM contributions to the nearby WISE--Star and the distant WISE--AGN sources are different; the distribution of source-intrinsic RMs of the WISE--AGNs is at most four times as wide as the distribution of intrinsic RMs of the star-forming WISE--Star galaxies if the distribution of intrinsic RMs of the WISE--Star sources itself is at least several rad m$^{-2}$ wide; and \vskip 0.2cm \noindent (6) we also detect no evolution of RM with redshift, suggesting that the RM is a product of the intrinsic properties of the radio galaxy and not a result of the intervening large-scale structure of the Universe. | 14 | 4 | 1404.1638 |
1404 | 1404.6353_arXiv.txt | {} {We study the star-formation history of the Galactic bulge, as derived from the age distribution of the central stars of planetary nebulae that belong to this stellar population. } { The high resolution imaging and spectroscopic observations of 31 compact planetary nebulae are used to derive their central star masses. We use the Bl\"ocker post asymptotic giant branch (post-AGB) evolutionary models, which are accelerated by a factor of three in this case to better fit the white dwarf mass distribution and asteroseismological masses. Initial-final mass relations (IFMR) are derived using white dwarfs in clusters. These are applied to determine original stellar masses and ages. The age distribution is corrected for observational bias as a function of stellar mass. We predict that there are about 2000 planetary nebulae in the bulge. } {The planetary nebula population points at a young bulge population with an extended star-formation history. The Bl\"ocker tracks with the cluster IFMR result in ages, which are unexpectedly young. We find that the Bl\"ocker post-AGB tracks need to be accelerated by a factor of three to fit the local white dwarf masses. This acceleration extends the age distribution. We adjust the IFMR as a free parameter to map the central star ages on the full age range of bulge stellar populations. This fit requires a steeper IFMR than the cluster relation. We find a star-formation rate in the Galactic bulge, which is approximately constant between 3 and 10\,Gyr ago. The result indicates that planetary nebulae are mainly associated with the younger and more metal-rich bulge populations. } {The constant rate of star-formation between 3 and 10\,Gyr agrees with suggestions that the metal-rich component of the bulge is formed during an extended process, such as a bar interaction.} | The origin and evolution of the Galactic bulge (GB) of the Milky Way is an area of active research. Bulges were believed to be similar to elliptical galaxies with a proposed origin in minor mergers. This was supported by the relation between bulge/spheroid luminosities and central black-hole masses \citep{G2007}. However, some bulges are more disk-like in their properties and these have been termed 'pseudo-bulges': their origin is likely unrelated to those of the spheroidal bulges \citep{KK2004}. Pseudo-bulges can form by secular evolution over a longer period, while classical bulges are predominantly an early product of galaxy formation. Whether the Milky Way bulge is a classical or pseudo-bulge is disputed. There is evidence for old stellar populations. \citet{V2009} compared OGLE and 2MASS data to the galactic model TRILEGAL to find good fits for models with a starburst from 8 to 10\,Gyr ago with the best fit for the younger age. The metallicity is close to solar with a distribution extending not very far towards lower values. In recent years, however, the evidence for a younger stellar component has been growing. In a recent review, \citet{Babu2012} concluded that the other characteristics point toward a mix of stellar populations although the main shape of the GB seems to be driven by secular evolution. This multi-component structure is also seen in the shape of the GB \citep{Robin2012}, which may be a combination of a classical and a pseudo (box-shape) bar. Regarding the age and duration of star-formation in GB, \citet{Tsuji2012} concluded that the metal-poor and metal-rich populations are formed at different times with the former extending over 2\,Gyr and the latter over 4\,Gyr. This separation was also found by \citet{Bensby2010, Bensby2011, Bensby2013}. They carried out spectroscopy of GB dwarf stars during gravitational lensing amplification events. Stellar parameters were derived based on LTE model atmospheres, and stellar ages were interpolated from a grid of isochrones. Bensby et al. found that stars with sub-solar metallicities are predominantly old (around 10\,Gyr) as expected in the GB. At super-solar metallicities, however, they find a wide range of ages with a peak around 5\,Gyr and a tail towards higher ages. They conclude that the origin of the GB is still poorly constrained and speculate that the young stars might be the manifestation of the inner thin disk. \citet{Gonzalez2011} also found evidence for a separate, high metallicity component of the bulge, which they attribute to stars originating from the thin disk. Planetary nebulae (PN), which are bright and easy to identify, can help us to address the question of the age of the Galactic bulge. The PNe trace populations which have ages ranging from $<1$ to 10\,Gyr. In contrast, red clump data analysis, as used by \citet{Robin2012}, cannot discriminate any ages over 5\,Gyr. However, deriving ages for the central stars of individual PNe has been a challenge. In this paper, we investigate the relation between the stellar populations and the planetary nebula population of the Galactic bulge by deriving the initial masses and, therefore, the ages of the PN central stars. This makes use of new Hubble Space telescope (HST) images and complementary Very Large Telescope (VLT) high-resolution spectra to find the mass distribution of the central stars of bulge planetary nebulae. These are converted to initial masses using initial-final mass relations (IFMR); the stellar ages are found using the stellar mass-age relations. The central star mass distribution can be compared with a well-known distribution of white dwarf masses and the age distribution can be compared with ages of bulge stars. The steps are critically dependent on post asymptotic giant branch (post-AGB) stellar evolution models and the IFMR. We show that a recalibration of the speed in the post-AGB evolution is required to reproduce the well-known distribution of white dwarf masses. The IFMR is used as a free parameter to fit known bulge ages. The final conversion to a star-formation history also requires corrections for observational bias. After these corrections, we find that the results are consistent with an extended epoch of bulge formation. The paper is organized as follows. Section\,\ref{obse_anal} presents the collected observational material and its analysis with photoionization models. In Sect.\,\ref{final_m}, masses of central stars of PNe are derived and compared to the known white dwarf masses. The discrepancies can be resolved with a proposed acceleration of post-AGB evolution. In Sect.\,\ref{ini_fin_m}, the stellar initial masses and ages from the zero-age main sequence (ZAMS) are derived using a newly-derived IFMR and some evolutionary model tracks. Section\,\ref{pn_popul} considers the effect of selection and PN visibility bias. Finally, the adjusted age distribution of the GB PNe is derived in Sect.\,\ref{sfh_b}, where we use the IFMR as a free parameter. When compared to other published bulge age determinations, it points towards a rather steep IFMR. | The purpose of the investigation is to use the masses and ages of the central stars of planetary nebulae to constrain their evolution and to derive a star-formation history of the Galactic bulge. Our conclusions are as follows: \begin{itemize} \item We use HST images and VLT echelle spectra of 31 compact bulge planetary nebulae to derive expansion ages for the nebulae. The Torun models are used to calculate the mass-averaged expansion velocities. A correction factor is derived from published hydrodynamical models \citep{Perinotto2004} to relate this to the expansion of the outer radius. The expansion age of the nebula and the temperature of the central star are fitted to post-AGB model tracks to derive masses of the central stars. \item We use the set of Bl\"ocker post-AGB tracks. Two problems are identified with these tracks: the extremely slow evolution of the lowest mass tracks prohibits the formation of an ionized planetary nebula, and the final stellar masses are higher than white dwarf masses and asteroseismology masses. To force consistency, we exclude the lowest mass model from the model interpolations. We also make the ad-hoc adjustment that the envelope masses could be overestimated to accelerate the Bl\"ocker tracks by a constant factor of three. This produces consistency with white dwarf masses and allows for the lowest masses of PN central stars found from asteroseismology. We note that the white dwarf mass distribution in the bulge is not known and that the acceleration proposed as constant can actually depend on stellar mass and envelope mass. \item An IFMR is applied to the central star masses to derive the original (ZAMS) masses. An observational relation is derived from white dwarf masses in clusters. This gives a relation in very good agreement with \citet{Casewell2009} but with a shallowing at the lowest masses to allow for 0.53\,$M_\sun$ white dwarf masses in globular clusters. A simple relation is proposed to estimate the total ages of the stars. \item The derived ages are lower than seems plausible for bulge populations. To improve the agreement between the derived ages and the published data on Galactic bulge ages, we use the IFMR as a free parameter and adjust it to a linear IFMR that relates our youngest and oldest object with corresponding youngest and oldest bulge stars in \citet{Bensby2013}. Such an IFMR appears steeper than other observational relations but is not as steep as some assumed for evolutionary calculations. With this IFMR, the total ages of our bulge objects cover the desired range. A further acceleration of the Bl\"ocker tracks is a possible alternative to adjusting IFMR. \item To convert the age distribution of the PN central stars to a star-formation history, the inherent bias needs to be quantified. We estimate the total number of PNe present in the bulge as 2000. We calculate the visibility time as a compact PN as a function of final mass and the stellar death rate (or PN birth rate) as function of age of the stellar population. The age distribution is corrected for these functions. After the correction an apparent peak at 3\,Gyr disappears, which confirms the role of observational bias. \item The age distribution obtained by us is consistent with the subsample of \citet{Bensby2013} with metallicities close to solar values. The derived star-formation history shows a constant rate over time in general. \item The set of proposed procedures was eventually verified by applying it in an inverted version to the final, approximately constant SFH. This resulted (as expected) in a core-mass distribution similar to the one obtained from PNe with an accuracy of about $0.01\,M_\sun$. The derived masses, ages and relations compose a rather consistent picture. \item Our results agree with an extended star formation in the Galactic bulge, which was approximately constant between 3 and 10\,Gyr ago. This is more consistent with a secular development of the bulge, as expected from interaction with a bar. There may still be an older component of the bulge with a different origin, as the metal-poor population in the bulge region seems to be poorly represented among the compact PNe. \end{itemize} | 14 | 4 | 1404.6353 |
1404 | 1404.1435_arXiv.txt | Time evolution of a black hole lattice universe with a positive cosmological constant $\Lambda$ is simulated. The vacuum Einstein equations are numerically solved in a cubic box with a black hole in the center. Periodic boundary conditions on all pairs of opposite faces are imposed. Configurations of marginally trapped surfaces are analyzed. We describe the time evolution of not only black hole horizons, but also cosmological horizons. Defining the effective scale factor by using the area of a surface of the cubic box, we compare it with that in the spatially flat dust dominated FLRW universe with the same value of $\Lambda$. It is found that the behaviour of the effective scale factor is well approximated by that in the FLRW universe. Our result suggests that local inhomogeneities do not significantly affect the global expansion law of the universe irrespective of the value of $\Lambda$. | \label{sec:intro} The so-called ``black hole lattice universe" has been firstly investigated by Lindquist and Wheeler in 1957\cite{RevModPhys.29.432}. They regularly arranged $N$ potions of the Schwarzschild spacetime on a virtual 3-sphere ($N=5$, 8, 16, 24, 120 and 600), and discussed evolution of this lattice universe based on the intuitively derived junction conditions between the Schwarzschild shell and the 3-sphere. The black hole lattice universe is often used as one of tools to evaluate effects of local non-linear inhomogeneities on the global expansion. Recently, black hole lattice universe models have been revisited by several authors\cite{Yoo:2013yea,Yoo:2012jz,Clifton:2012qh,Clifton:2009jw,Uzan:2010nw,Bentivegna:2012ei,Bentivegna:2013xna,Bruneton:2012cg,Clifton:2013jpa,Clifton:2014lha,Korzynski:2013tea}. Time symmetric initial data for $N$-black hole systems on a virtual 3-sphere have been analyzed in Refs.~\cite{Clifton:2012qh,Korzynski:2013tea}. Time evolution of the 8-black hole system has been performed and analyzed in Ref.~\cite{Bentivegna:2012ei}. Initial data for a black hole inside a cubic box with a periodic boundary condition have been constructed and analyzed in Refs.~\cite{Yoo:2012jz,Bentivegna:2013xna}, and those time evolutions have been investigated in Refs.~\cite{Yoo:2013yea,Bentivegna:2012ei}. We call this cubic lattice model the ``black hole universe" in this paper. The purpose of this paper is to extend the black hole universe so that it admits a positive cosmological constant. In Ref.~\cite{Yoo:2013yea}, it has been reported that, if the box size of the black hole universe is sufficiently larger than the horizon radius, the global expansion law can be well approximated by that in the Einstein-de Sitter universe. Our final purpose is to check this fact with a positive cosmological constant. Since our universe is likely to be filled with dark energy components, such as the positive cosmological constant, it is important to investigate the effect of local non-linear inhomogeneities on the global expansion law with the cosmological constant. In the all references listed above, the cosmological constant is set to be zero. Therefore, solving technical problems to consider non-zero cosmological constant cases, we investigate it in this paper. One of non-trivial technical problems is how to construct an initial data set which is appropriate as a initial condition for the time evolution. In this paper, we describe a procedure to construct puncture initial data for the black hole universe with a positive cosmological constant. Another interesting problem is to find different kinds of marginal surfaces. As in the case of Kottler(Schwarzschild-de Sitter) solution, the black hole universe with a positive cosmological constant can have not only black hole horizons but also de Sitter cosmological horizons. As far as we know, it is first time to numerically find the ${\bf S}^2$ cosmological horizons without any symmetry which makes it possible to reduce the number of the effective dimension. To check the existence and structure of marginal surfaces is very useful to understand the spacetime structure. This paper is organized as follows. In Sec.~\ref{sec:initial}, we describe how to construct initial data of the black hole universe with a positive cosmological constant. Then, we analyze the structure of the initial data in Sec.~\ref{sec:marginal} searching for different kinds of marginal surfaces. In Sec.~\ref{sec:evolve}, time evolutions are described. The evolution of the configuration of marginal surfaces and the expansion law are discussed there. Sec.~\ref{sec:summary} is devoted to a summary. In this paper, we use the geometrized units in which the speed of light and Newton's gravitational constant are one, respectively. | \label{sec:summary} In this work, a black hole lattice universe model with positive cosmological constant has been simulated. The construction of puncture initial data with a positive cosmological constant has been described in Sec.~\ref{sec:initial}. The vacuum Einstein equations in a cubic box with a black hole in the center have been numerically solved with periodic boundary conditions by using the BSSN formalism\cite{Shibata:1995we,Baumgarte:1998te}. Configurations of marginal surfaces on the initial hypersurfaces and those time evolution have been analyzed. We found two impressive transitions of the configuration in time evolution: appearance of the outer cosmological horizon and the bifurcation surface crossing. Finally, comparing the effective scale factor defined by the surface area and the scale factor for the corresponding flat dust FLRW universe, we have concluded that the expansion law of the black hole universe can be well approximated by that of the corresponding flat dust FLRW universe in the sufficiently late time irrespective of the value of the cosmological constant. | 14 | 4 | 1404.1435 |
1404 | 1404.4817_arXiv.txt | The most iron-poor stars in the Milky Way provide important observational clues to the astrophysical objects that enriched the primordial gas with heavy elements. Among them, the recently discovered iron-deficient star SMSS J031300.36-670839.3 shows a remarkable chemical composition with non-detection of iron ([Fe/H]$<-7.1$) and large enhancement of carbon and magnesium relative to calcium. We investigate supernova yields of metal-free (Population III) stars to interpret the abundance pattern observed in this star. We report that the high [C/Ca] and [C/Mg] ratios and upper limits of other elemental abundances are well reproduced with the yields of core-collapse supernovae (that have normal kinetic energies of explosion $E$ of $E_{51}=E/10^{51}$erg$=1$) and hypernovae ($E_{51}\geq 10$) of Population III 25$M_{\odot}$ or 40$M_{\odot}$ stars. The best-fit models assume that the explosions undergo extensive matter mixing and fallback, leaving behind a black hole remnant. In these models, Ca is produced by static/explosive O burning and incomplete Si burning in the Population III supernova/hypernova, in contrast to the suggestion that Ca is originated from the hot-CNO cycle during the presupernova evolution. Chemical abundances of four carbon-rich iron-poor stars with [Fe/H]$<-4.5$, including SMSS J031300.36-670839.3 are consistently explained by the faint supernova models with the ejected mass of $^{56}$Ni less than 10$^{-3}M_{\odot}$. | \label{sec:intro} Characteristic masses of the first stars (Population III or Pop III stars) and the nature of their supernova explosions are critically important to determine their role in the cosmic reionization and subsequent star formation in the early universe \citep[e.g.][]{bromm11}. Cosmological simulations have shown that the Pop III stars could be very massive $\gtrsim 100M_{\odot}$, as the result of cooling of primordial gas via hydrogen molecules \citep [e.g.][]{bromm04}. More recent studies, however, propose the mechanisms in which lower mass stars can form through radiation feedback from growing protostars and/or disk fragmentation \citep{hosokawa11,clark11,hirano13,bromm13,susa13}. Abundance patterns of the lowest-metallicity stars in our Galaxy provide us with a rare opportunity to observationally constrain the masses of the Pop III stars. The chemical abundances in the four iron-poor stars with [Fe/H]$<-4.5$, HE 0107-5240 \citep{christlieb02}, HE 1327-2326 \citep{frebel05,aoki06}, HE 0557-4840 \citep{norris07}, and SDSS J102915$+$172927 \citep{caffau11} (see \citet{hansen14} for a recent discovery of another metal-poor star in this metallicity range) do not show signature of pair-instability supernovae of very massive ($\gtrsim 140M_{\odot}$) stars as their progenitors \citep[][and references therein]{nomoto13}. Instead, the observed abundances in these stars are better explained by the yields of core-collapse supernovae of moderately massive Pop III stars with several tens of $M_{\odot}$ \citep{umeda02,umeda03,limongi03,iwamoto05,tominaga07b, tominaga09,tominaga14, heger10, kobayashi14}. Another important insight into the nature of the Pop III stars is that a large fraction of most iron-poor stars are carbon-rich \citep[e.g.][]{hansen14}. \citet{iwamoto05} suggests that the large enhancement of carbon observed in both HE 0107-5240 and HE 1327-2326 is explained by their models with mixing of supernova ejecta and their subsequent fallback on to the central remnant. These models with different extent of the mixing regions simultaneously reproduce the observed similarity in [C/Fe] and more than a factor of $\sim 10$ differences in [O, Mg, Al/Fe] between the two stars. A metal-poor star SMSS J031300.36-670839.3 (SMSS J0313-6708), recently discovered by SkyMapper Southern Sky Survey, provides us with a new opportunity to test theoretical predictions about the Pop III stars \citep{keller07,keller14}. Follow-up spectroscopic observations found that this object shows an extremely low upper limit for its iron abundance ([Fe/H]$<-7.1$); more than $1.0$ dex lower than the previous record of the lowest iron-abundance stars. In this letter, we extend the study of \citet{iwamoto05} and examine whether the abundances of the five stars with [Fe/H]$<-4.5$, including the most iron-deficient star SMSS 0313-6708, can be consistently explained by the supernova yields of Pop III stars which undergo the mixing and fallback. | \label{sec:discussion} As shown in the previous section, the abundance measurements (C, Mg, and Ca) in SMSS 0313-6707 are reproduced with the Pop III supernova/hypernova yields ($(M, E_{51})=(25, 1), (25, 10), (40, 1)$ and $(40, 30)$), with the model parameters corresponding to faint supernova/hypernova with extensive mixing and prominent fallback. In order to discriminate models with different masses and energies, additional abundance measurements for oxygen as well as iron-peak elements including V, Mn, Co, and Cu are particularly useful. The ejected mass of $^{40}$Ca is only $\sim 10^{-7}-10^{-8} M_{\odot}$ in the faint supernovae/hypernovae models. To be compatible with the observed calcium abundance ([Ca/H]$=-7.0$), the supernova ejecta should be diluted with $\sim 10^{3}- 10^{4}M_{\odot}$ of the primordial gas. In the case of the supernova models with $E_{51}=1$, this is consistent with the suggested relation between the supernova energy versus swept-up gas mass with primordial composition \citep{thornton98}, as adopted in \citet{tominaga07b}, for the assumed number density of hydrogen $1<n_{H}<100$ cm$^{-3}$. On the other hand, this relation predicts that the hypernova sweeps up much larger amount of hydrogen ($\gtrsim 10^{5}M_{\odot}$ for $n_{H}<10^4$ cm$^{-3}$) than the above values. A recent cosmological simulations of the transport of the metals synthesized in a Pop III supernova, however, suggests a wide range of metal abundances in the interstellar gas clouds after an explosion of a Pop III star \citep{ritter12}, and thus the hypernova could work as the source of the chemical enrichment for the formation of stars with [Ca/H]$\lesssim -7$. In our model, Ca is produced by hydrostatic/explosive O burning and incomplete Si burning in the Pop III supernova or hypernova with masses $25$ or $40 M_{\odot}$. This is different from the 60$M_{\odot}$ model adopted in \citet{keller14}, where Ca is originated from the hot CNO cycle during the main-sequence phase. Synthesis of $\sim 10^{-7}M_{\odot}$ Ca in the hot CNO cycle is also seen in the $100 M_{\odot}$ models of \citet{umeda05}. On the other hand, Ca produced in this mechanism is not seen in the 25 and 40$M_{\odot}$ progenitors analyzed in this paper. The mass fraction of Ca near the bottom of the hydrogen layer in these progenitors is $\log X_{\rm Ca}<-10$. In order to clarify which of these nucleosynthesis sites are responsible for the observed Ca, we note a different prediction between the two scenarios. Our models suggest that a certain amount of Fe distributed in the inner region as well as explosively synthesized Ca are ejected as a result of the assumed mixing at $M_{r}=2-6M_{\odot}$. This results in the [Fe/Ca] ratio of $\sim -2$ - $0$, depending on the $M_{\rm cut}({\rm ini})$. Consequently, our models of the faint supernova/hypernova predict the metallicity distribution function (MDF) to be continuous even below [Fe/H]$<-6$. On the other hand, the model adopted in \citet{keller14}, in which Ca is produced in the hot-CNO cycle, predicts [Fe/Ca]$\lesssim-3$, which is not observed in other extremely iron-poor stars. Because of the distinct Ca production sites, the MDF could be discontinuous in the most metal-poor region. Future photometric and spectroscopic surveys to discover lowest metallicity stars and their MDF provide clues to discriminate these mechanisms. The models adopted in this work suggest that the faint Pop III supernovae could be the origin of the observed abundance patterns and the variation among the most iron-poor carbon-rich stars. To understand physics of faint Pop III supernovae, multi-dimensional simulations are necessary. A large-scale mixing as suggested for the carbon-enhanced stars are not predicted in the models with Rayleigh-Taylor mixing alone \citep{joggerst09}. Instead, a more likely origin of such large-scale mixing-fallback would be the jet-induced supernova/hypernova, where the inner material can be ejected along the jet-axis while a large fraction of the material along the equatorial plain falls back \citep{tominaga09}. | 14 | 4 | 1404.4817 |
1404 | 1404.0433_arXiv.txt | { Ultracompact dark matter minihalos (UCMHs) would be formed during the earlier universe if there were large density perturbations. If the dark matter can decay into the standard model particles, such as neutrinos, these objects would become the potential astrophysical sources and could be detected by the related instruments, such as IceCube. In this paper, we investigate the neutrino signals from the nearby UCMHs due to the gravitino dark matter decay and compare these signals with the background neutrino flux which is mainly from the atmosphere to get the constraints on the abundance of UCMHs. | % \label{sect:intro} Structure formation is one of important research fields in cosmology. According to the theory, the present cosmic structures originated from the earlier density perturbations with the amplitude $\delta\rho/\rho\sim10^{-5}$ and this has been confirmed by many observations. On the other hand, primordial black holes would be formed if there were large density perturbations ($\delta\rho/\rho>0.3$) in the earlier universe (\citealt{PBH}). Recently, \cite{Ricott+Gould} found that if the amplitude of density perturbations was between above values a new kind of dark matter structures named ultracompact dark matter minihalos (UCMHs) would be formed. Compared with the classical dark matter halos, the formation time of these objects is earlier ($z\sim 1000$) and the density profile is steeper ($\rho(r) \sim r^{-2.25}$). If dark matter is in the form of weakly interacting massive particles (WIMPs), such as the neutralino, they can annihilate into the standard model particles, such as photons, positrons or neutrinos (\citealt{dm_1,dm_2}). Moreover, because the dark matter annihilation rate is proportional to the square of number density, the UCMHs would become one kind of the potential astrophysical sources (\citealt{scott_prl,scott_prd,prd_positron,prd_neutrino}). Besides the annihilation, decay is another important approach to detect dark matter signals. This is especially crucial for those dark matter candidates which do not annihilate. A famous example is the gravitino dark matter which in some supergravity models is the lightest supersymmetric particle (\citealt{dm_2}). Although compared with the annihilation the decay rate is proportional to the number density instead of the number density square, the decay is still very important for the cosmological probes for the dark matter particles which do not have the annihilation channels. \cite{epl} have investigated the gamma-ray flux from nearby UCMHs due to the dark matter decay. Through comparing with the observations they obtained the constraints on the abundance of UCMHs for different decay channels, lifetimes and density profiles of Milky Way. % They found that the strongest constraint comes from the $b\overline{b}$ channel with the dark matter mass $m_{\chi}=100\mathrm{GeV}$, the fraction of UCMHs is $f_{\mathrm{UCMHs}}\sim5\times10^{-5}$. Besides the high energy photons the other kind of important products of dark matter decay are neutrinos and they usually accompany with photons. The advantage of neutrino detection is that neutrinos can propagate in the space without attenuation due to its very weakly interaction with other particles. Therefore, comparing with other particles (e.g. electrons and positrons) the orientation of corresponding sources can be confirmed directly. When neutrinos propagate through the medium, such as the ice, muons ($\mu$) can be produced by the charged current interaction and detected by the Cherenkov radiation detector. Because the neutrino signals accompanying with the production of gamma-ray, the study on neutrino signals would be complementary to the gamma-ray observations especially for the larger dark matter mass and lepton channels (\citealt{com_neu_1,com_neu_2,com_neu_3}). In this paper, we will investigate the neutrino signals from UCMHs due to the dark matter decay. As we have not observed any excess of neutrino flux comparing with the atmospheric neutrino flux, we get the constraints on the abundance of UCMHs. This paper is organised as follows. The neutrino flux from nearby UCMHs due to dark matter decay are studied in section 2. In section 3, we obtain the constraints on the fraction of UCMHs. We conclude with discussions in section 4. | The UCMHs would be formed if there are large density perturbations during the earlier epoch and then its cosmological abundance become one of the important issuses. In the previous works, the main limits are from the research of the gamma-ray flux due to dark matter annihilation within UCMHs. In our recently works, we considered the constraints from the neutrino flux due to the dark matter annihilation and the gamma-ray flux due to the dark matter decay (\cite{epl,prd_neutrino}). In this work, we extended these works and investigated the neutrino flux from nearby UCMHs due to the gravitino particles decay. The decay styles of gravitino include two-body and three-body decays. The latter can also provide one of the explanations of the positrons (or positrons plus electrons) excess which have been observed recently by the PAMALE and Fermi. In this work, we mainly considered the three-body decay style. Most of this decay productions are leptons, therefore, the neutrinos will be plenty. We researched the neutrino flux from nearby UCMHs due to the dark matter decay and compared that with the signals from the atmosphere which is the main background of neutrino detection. We found that for the larger dark matter mass or the UCMHs the final flux would excess the ATM. These results are similar to the dark matter annihilation cases (\citealt{prd_neutrino}). On the other hand, because we have not observed any excess of neutrino flux from nearby unknown sources, so the abundance of UCMHs can be constrained. We considered ten years exposure times for neutrino observation and signals with 2$\sigma$ statistic significance to get the limits on the fraction of UCMHs. In this work, we also assumed that the distribution of these objects is uniform in the universe. We found that the strongest limits on the abundance of UCMHs is $f_\mathrm{UCMHs} \sim 10^{-3}$. One should note that these results depend on the dark matter mass. From the Fig.~\ref{fig:cons}, it can be seen that the limits will be stronger for the larger dark matter mass. Moreover, the final constraints are stronger for the upward events comparing with the contained events. The other important parameters is the dark matter decay rate. In this paper, we have used the conservative value, $\Gamma = 10^{-26} \mathrm{s^{-1}}$ for our research. There are many works where the constraints on the decay rate of gravitino are obtained. \cite{Huang:2011xr} used the Fermi observations of nearby galaxy clusters to get the constraints on this parameter. They found that the limits on the lifetimes of gravitino from the clusters observations is $\tau(1/\Gamma) \sim 2 \times 10^{26}s$ and it is mostly independent of on the dark matter mass. The limits obtained from the lines signals are different and the lifetime becomes smaller with the increasing of dark matter mass. For the future neutrino observations, such as KM3Net, because its effective area and volume will be improved, it is expected that the constraints on the abundance of UCMHs will be much more stronger. \label{sect:discussion} \normalem | 14 | 4 | 1404.0433 |
1404 | 1404.7025_arXiv.txt | The IceCube observation of cosmic neutrinos with $E_{\nu} > 60$ TeV, most of which are likely of extragalactic origin, allows one to severely constrain Lorentz invariance violation (LIV) in the neutrino sector, allowing for the possible existence of superluminal neutrinos. The subsequent neutrino energy loss by vacuum $e^+e^-$ pair emission (VPE) is strongly dependent on the strength of LIV. In this paper we explore the physics and cosmology of superluminal neutrino propagation. We consider a conservative scenario for the redshift distribution of neutrino sources. Then by propagating a generic neutrino spectrum, using Monte Carlo techniques to take account of energy losses from both VPE and redshifting, we obtain the best present constraints on LIV parameters involving neutrinos. We find that $\delta_{\nu e } = \delta_{\nu} - \delta_e \le 5.2 \times 10^{-21}$. Taking $\delta_e \le 5 \times 10^{-21}$, we then obtain an upper limit on the superluminal velocity fraction for neutrinos alone of $1.0 \times 10^{-20}$. Interestingly, by taking $\delta_{\nu e} = 5.2 \times 10^{-21}$, we obtain a cutoff in the predicted neutrino spectrum above 2 PeV that is consistent with the lack of observed neutrinos at those energies, and particularly at the Glashow resonance energy of 6.3 PeV. Thus, such a cutoff could be the result of neutrinos being slightly superluminal, with $\delta_{\nu}$ being $(0.5 \ {\rm to} \ 1.0) \times 10^{-20}$. | The possible existence of superluminal neutrinos as a consequence of Lorentz invariance violation (LIV) was brought to the attention of the physics community by their apparent observation ~\cite{ad11}. Shortly thereafter, Cohen and Glashow~\cite{cg11} presented a powerful theoretical argument against the results in Ref.~\cite{ad11}. Their argument was based on the implication that these neutrinos would rapidly lose energy by the dominant energy loss channel of vacuum electron-positron pair emission (VPE), i.e., $\nu \to \nu \,e^+\, e^-$. Eventually, the results in Ref.~\cite{ad11} were retracted~\cite{ad12}. (See also Ref.~\cite{ad13}). The Ice Cube collaboration has recently reported the observation of 37 extraterrestrial neutrinos with energy above $\sim$ 60 TeV, giving a cosmic neutrino signal $5.7\sigma$ above the atmospheric background~\cite{wh14}. This is significant evidence for a neutrino flux of cosmic origin, above that produced by atmospheric cosmic-ray secondaries~\cite{aa13a}.% The very existence of PeV neutrinos has been used to place strong constraints on LIV in the neutrino sector ~\cite{st14},~\cite{bo13}.% | \begin{figure}[t] {\includegraphics[scale=0.6]{spectra.eps}} \caption{Calculated neutrino spectra with VPE and redshifting compared with the IceCube data both including a subtraction of atmospheric charm $\nu$'s at the 90\% C.L. (cyan) and omitting such a subtraction (black)~\cite{wh14}. Curves from left to right are spectra obtained with rest-frame threshold energies of 1, 2, 4, 10, 20 and 40 PeV. The corresponding values of $\delta_{\nu e}$ are given by equation (\ref{spectra}).} \label{spectra} \end{figure} The results of our calculations show that there is a high-energy drop off in the propagated neutrino spectrum resulting from the opening of the VPE channel above threshold. Furthermore, the redshifting effect pushes the cutoff in the energy spectrum below the non-redshifted rest-frame threshold energy. As discussed before, we assume that the neutrino production rate follows the star formation rate in redshift space. This rate peaks at a redshift between 1 and 2. The neutrinos emitted during this past era of enhanced stellar and galactic activity are then redshifted by a factor of 2 to 3. The redshifting effect dominates the shape of the resulting spectra regardless of threshold energy. This is because the mean propagation time is very short compared with the total travel time with the exception of rest-energy thresholds greater than 10 PeV as follows from equations (\ref{threshold}) - (\ref{redshift}) (See table~\ref{proptable}). Furthermore, the mean propagation time is also short for {\em all} energies greater than the threshold, with the exception of only those very near threshold, as illustrated in Figure \ref{proptime} for a rest-energy threshold of 10 PeV. In the case of rest-energy thresholds greater than 10 PeV, the particles very near threshold will simply redshift below it without decay. This has little impact on their final observed energies at $z = 0$. Our calculated neutrino spectra follow our assumed $E^{-2}$ power-law form below $\sim$0.2 of the the redshifted VPE threshold, have a small pileup effect up to the redshifted threshold energy, and have a sharp high energy cutoff at higher energies, as shown in Figure~\ref{spectra}. The pileup is caused by the propagation of the higher energy neutrinos in energy space down to energies within a factor of $\sim$5 below the threshold. This is indicative the fact that fractional energy loss from the last allowed neutrino decay before the VPE process ceases is 0.78~\cite{cg11}. The pileup effect is similar to that of energy propagation for ultrahigh energy protons near the GZK threshold~\cite{st89}. Our results yield the best constraints LIV in the neutrino sector to date, {\it viz.}, $\delta_{\nu e} = \delta_{\nu} - \delta_e \le \ 5.2 \times 10^{-21}$. This is because our results for our rest-frame threshold energy cases below 10 PeV as shown in Figure~\ref{spectra} are inconsistent with the IceCube data. Our result for a 10 PeV non-redshifted threshold, corresponding to $\delta_{\nu e} = \ 5.2 \times 10^{-21}$, is just consistent with the IceCube results, giving a cutoff effect above 2 PeV. We note that the present best upper limit on $\delta_e$ is $5 \times 10^{-21}$~\cite{st14}. Thus for the conservative case of no-LIV effect, {\it e.g.}, if one assumes a cutoff in the intrinsic neutrino spectrum of the sources or one assumes a steeper assumed PeV neutrino spectrum proportional to $E_{\nu}^{-2.3}$~\cite{an13,wh14}, we find the new constraint on superluminal neutrino velocity, $\delta_{\nu} = \delta_{\nu e} + \delta_e \le \ 1.0 \times 10^{-20}$. However, the steeper spectrum scenario has been placed into question~\cite{ch14}. Interestingly, for an $E_{\nu}^{-2}$ power-law neutrino spectrum, we find the possibility that the apparent cutoff in the observed spectrum above $\sim$2 PeV can conceivably be an effect of Lorentz invariance violation (see Figure~\ref{spectra}). (Another suggestion involving LIV effects of {\it subluminal} neutrinos has recently been discussed~\cite{an14}). A hard $E_{\nu}^{-2}$ spectrum has been proposed to be produced in starburst galaxies~\cite{lo06}. The IceCube flux is below the upper limit of $2 \times 10^{-8} \ {\rm GeV}{\rm cm}^{-2}{\rm s}^{-1}{\rm sr}^{-1}$ obtained by one of us for the neutrino flux from starburst galaxies~\cite{st07}, allowing for this possibility. The power-law source spectrum option opens the possibility that a high energy cutoff in such a hard $E_{\nu}^{-2}$ power-law neutrino spectrum could be caused by a small violation of Lorentz invariance, with neutrinos being very slightly superluminal, with $\delta_{\nu}$ being $(0.5 \ {\rm to} \ 1.0) \times 10^{-20}$, taking $0 \le \delta_e \le 0.5 \times 10^{-20}$. As has been pointed out previously for ultrahigh energy neutrinos~\cite{go12}, one test for the cutoff scenario would be the non-observation of the "cosmogenic" neutrinos from photopion production interactions of ultrahigh energy cosmic rays with the cosmic background radiation~\cite{be69}, since all cosmological neutrinos above $\sim$2 PeV would be affected by the VPE process. Such a non-observation would have implications for $\gamma$-ray constraints on ultrahigh energy cosmic ray origin and composition models, perhaps implying the ultrahigh energy cosmic rays are mainly heavy nuclei~\cite{al12}. | 14 | 4 | 1404.7025 |
1404 | 1404.5957_arXiv.txt | {The damping of a non-uniform magnetic field between the redshifts of about $10^4$ and $10^6$ injects energy into the photon-baryon plasma and causes the CMB to deviate from a perfect blackbody spectrum, producing a so-called $\mu$-distortion. We can calculate the correlation $\langle\mu T\rangle$ of this distortion with the temperature anisotropy $T$ of the CMB to search for a correlation $\langle B^2\zeta\rangle$ between the magnetic field $B$ and the curvature perturbation $\zeta$; knowing the $\langle B^2\zeta\rangle$ correlation would help us distinguish between different models of magnetogenesis. Since the perturbations which produce the $\mu$-distortion will be much smaller scale than the relevant density perturbations, the observation of this correlation is sensitive to the squeezed limit of $\langle B^2\zeta\rangle$, which is naturally parameterized by $b_{\text{NL}}$ (a parameter defined analogously to $f_{\text{NL}}$). We find that a PIXIE-like CMB experiments has a signal to noise $S/N\approx 1.0 \times b_{\text{NL}} (\tilde B_\mu/10\text{ nG})^2$, where $\tilde B_\mu$ is the magnetic field's strength on $\mu$-distortion scales normalized to today's redshift; thus, a 10 nG field would be detectable with $b_{\text{NL}}=\mathcal{O}(1)$. However, if the field is of inflationary origin, we generically expect it to be accompanied by a curvature bispectrum $\langle\zeta^3\rangle$ induced by the magnetic field. For sufficiently small magnetic fields, the signal $\langle B^2 \zeta\rangle$ will dominate, but for $\tilde B_\mu\gtrsim 1$ nG, one would have to consider the specifics of the inflationary magnetogenesis model. We also discuss the potential post-magnetogenesis sources of a $\langle B^2\zeta\rangle$ correlation and explain why there will be no contribution from the evolution of the magnetic field in response to the curvature perturbation.} | \label{sec:introduction} The $\Lambda$CDM model of cosmology provides a very good fit to many cosmological observables. However, there remain many questions about the evolution of the early universe and the forces that shaped it. Over the last decade, much of our cosmological information has come from exploring the temperature perturbations of the mostly uniform Cosmic Microwave Background (CMB) by experiments such as WMAP, Planck, SPT, and ACT. However, at this point, we have nearly exhausted the information in this signal (at least from pre-recombination sources), and future discoveries must come from other sources, e.g. large scale structure, improved standard candle measurements, the 21 cm line, etc. However, the temperature perturbation is not the only information in the CMB, as excitingly suggested by the possible recent $B$-mode detection of BICEP2 \cite{Ade:2014xna}, with other experiments soon to come with more polarization data, including Planck, Spider, the Keck Array, and POLARBEAR. In this paper, we look at still another source of data in the CMB: distortion from a perfect blackbody spectrum, in particular $\mu$-type distortion. To understand $\mu$-distortion, one must first realize the limits of a widely known fact: the CMB has no measured pre-recombination deviations from a blackbody spectrum (\cite{Fixsen:1996nj,Mather:1993ij,Mather:1991pc}). However, even at recombination, the photon distribution is actually expected to be only an imperfect Planck spectrum \cite{Weymann:1966ab,Zeldovich:1969ab,Sunyaev:1970aq}. In an ideal photon gas, both the energy distribution and number density of photons are specified by the photon temperature $T$; thus, adding energy or entropy to the gas generically requires both a redistribution of photon energy and a change in photon number in order to re-reach a Planck spectrum at some new temperature $T'$. Well before recombination, at $z\gg z_\mu^i =2\sci{6}$, the photon spectrum in the early universe plasma is Planckian because of the efficiency of the various relevant reactions. However for $z\lesssim z_\mu^i$, the primary photon non-conserving process, double-Compton scattering ($\gamma+e^-\to 2\gamma+e^-$), becomes inefficient, while elastic Compton scattering remains efficient; thus, the photon gas responds to entropy increases by reaching thermal equilibrium but with a conserved photon number, which is thermodynamically equivalent to giving the photons a chemical potential $\mu$. The result is that the photon develop a Bose-Einstein distribution given by the appropriate values of $T$ and $\mu$; this deviation is called $\mu$-distortion. Later at $z\lesssim z_\mu^f= 5\sci{4}$, even elastic Compton scattering is inefficient and new perturbations to the photon distribution are minimally reprocessed, usually producing so-called $y$-distortions. $\mu$-distortion is typically parameterized as a dimensionless number in the photon distribution function $n(x) = (e^{x+\mu}-1)^{-1}$, where $x$ is the dimensionless frequency $x\equiv h\nu/k_B T$. COBE/FIRAS constrained $|\mu|< 9\sci{-5}$ \cite{Fixsen:1996nj}, which was marginally improved by the TRIS experiment to $|\mu|< 6 \sci{-5}$ \cite{Zannoni:2008xx,Gervasi:2008eb}. The proposed PIXIE experiment would considerably improve the bound to $|\mu|\lesssim 9\sci{-8}$ \cite{Kogut:2011xw}. $\mu$-distortion is produced by effects that inject energy/entropy into the photon-baryon plasma during the $\mu$-distortion era $z_\mu^f < z < z_\mu^i$. There is one inevitable source of such injections, namely Silk damping of density perturbations \cite{Silk:1968,Peebles-Yu:1970,Kaiser:1983}. Particle decay during the appropriate era can also produce $\mu$-distortion \cite{Hu-Silk-Relic-pcls:1993}, though the decay time must be fine-tuned and one must take care to satisfy constraints on the radiation energy density from BBN and the CMB. In this paper, we will focus on the production of $\mu$-distortion from the decay of non-uniform magnetic fields due the viscosity of the photon-baryon plasma \cite{Jedamzik:1999bm}. Most $\mu$-distortion studies have focused on the monopole $|\mu|$, i.e. the sky-averaged $\mu$-distortion. Recently, \cite{Pajer&Zald:2012-New-window}\footnote{See also \cite{Ganc:2012ae}.} looked at the signatures of anisotropic $\mu$-distortion. In their case, they considered the $\mu$-distortion caused by the damping of the density perturbation $\zeta$. Since the entropy for the $\mu$-distortion comes from spectral power (or equivalently, from the energy in the density waves), one has that $\mu\propto \zeta^2$. They then correlated the $\mu$-distortion signal with the larger temperature anisotropy signal $T$, which is primarily given by $T\propto\zeta$. Thus, the $\braket{\mu T}$ correlation constrains the squeezed limit of the bispectrum $\braket{\zeta^3}$, i.e. $f_{\text{NL}}^\text{loc}$. As noted earlier, the damping of magnetic fields also produce $\mu$-distortion \cite{Jedamzik:1996wp,Jedamzik:1999bm}, proportional to the damped energy $B^2$. It is probable that there are $\sim\mu$G magnetic fields in galaxies and clusters (e.g. \cite{Bonafede:2010xg,Beck:2012ag}), and their origins are still speculative. More recently, there have been tantalizing hints of large-scale magnetic fields in the intergalactic medium with strengths of $\mcl{O}(10^{-15} - 10^{-20})$ G \cite{Tavecchio:2010mk,Neronov:1900zz,Takahashi:2013uoa} and which are incompatible with most magnetogenesis scenarios (e.g. from the QCD or electroweak phase-transition). Inflation can, in principle, produce such large-scale fields, though taking the BICEP2 data \cite{Ade:2014xna} at face values strongly disfavors inflationary production as well \cite{Kandus:2010nw,Ferreira:2013sqa,Fujita:2014sna,Ferreira:2014af}. Just as with the curvature perturbation, we can study cosmological magnetic fields through their correlations and indeed, to study magnetic fields, it makes sense to consider a possible correlation $\bbzet$ with the already-detectable curvature perturbation signal. In analogy with $f_{\text{NL}}$, we can suppose that a correlated magnetic field $\bB^\text{corr}$ is produced from an uncorrelated field $\bB^\text{uncor}$ \cite{Jain:2012ga} as \begin{align} \bB^\text{corr} \simeq \bB^\text{uncor} + b_{\text{NL}} \bB^\text{uncor} \zeta + \ldots\,, \end{align} so that \begin{align}\label{eq:dfnn-of-bnl} \left\langle\bB^*(\bk_1) \cdot \bB(\bk_2)\, \zeta(\bk_3) \right\rangle_{k_3\ll k_1\approx k_2} = (2\pi)^3 b_{\text{NL}} \delta^{(3)}\!(-\bk_1 + \bk_2 + \bk_3) P_B(k_1) P_\zeta(k_3) \end{align} in the squeezed limit $k_3\ll k_1\approx k_2$. Actually, we take (\ref{eq:dfnn-of-bnl}) to be the definition of $b_{\text{NL}}$, with the understanding that we are primarily interested in the squeezed limit. Generically, one expects that inflationary magnetogenesis produces\footnote{A consistency relation for a simple class of models was derived in \cite{Jain:2012ga,Jain:2012vm}, yielding $b_{NL}=(n_B-4)$.} $|b_{\text{NL}}|\gtrsim 1$. The $\braket{\mu T}$ correlation first explored by \cite{Pajer&Zald:2012-A-hydrodynamical} offers a promising way to probe such a correlation, which we explore here. Related ideas of how to probe $b_{NL}$ with large-scale structure consistency relations was put forward in \cite{Berger:2014wta}. In this work, we demonstrate that a primordial $\bbzet$ correlation\footnote{Note that, by ``primordial'', we mean existing in the era of magnetogenesis, not (as is sometimes meant) merely happening anytime before recombination.} could potentially be observable in a measurement of the $\braket{\mu T}$ correlation if $b_{\text{NL}}=\sO(1)$ and the magnetic field on $\mu$ scales is $\sO(10\text{ nG})$. However, one also needs to consider that magnetic field energy density is inherently non-Gaussian (since it goes as $B^2$) so that magnetic fields inevitably source some level of the three-point function $\langle\zeta^3\rangle$, which also generates a $\braket{\mu T}$ signal. If the magnetic fields were of inflationary origin, there would be sufficient time for this correlation to grow large. For sufficiently small magnetic fields, the $\bbzet$ signal will still dominate because it goes as $B^2$ whereas other signals go as higher powers of the magnetic field. However, for $B_\mu\gtrsim$ 1 nG, one has to consider the specifics of the inflationary magnetogenesis model to determine which signal would be detected first. Our work is organized as follows. In Sec. \ref{sec:measuring-mgntc-flds-mu-distortion}, we discuss how magnetic fields damp in the early universe, producing $\mu$-distortion; we then correlate this with the temperature perturbation $T$ to find a fairly generic formula for $C_l^{\mu T}$ in terms of $\bbzet$ without regard to the source of the correlation. In the next section, Sec. \ref{sec:primordial-bnl}, we quickly derive the formula for the signal from a primordial $b_{\text{NL}}$. In the next two sections, we consider two other sources for a $\bbzet$ correlation. In Sec. \ref{sec:evln-of-zet-due-to-B}, we repeat \cite{Miyamoto:2013oua}'s recent calculation of the correlation produced by the evolution of $\zeta$ due to the anisotropic stress of the magnetic field; they found (and we confirm) a very small signal. In Sec. \ref{sec:evln-of-B-due-to-zet}, we consider the reverse effect, the potential correlation caused by the evolution of $B$ due to the initial adiabatic perturbation, finding also no relevant impact. We discuss the observational impact of our results in Sec. \ref{sec:observ-prim-BBzet}, including the consequences of the competing $f_{\text{NL}}$ signal, and then conclude in Sec. \ref{sec:conclusion}. \subsection{Notation and conventions} \label{sec:notation} We use the following conventions: \begin{itemize} \item We use $t$ for physical time and $\eta$ for conformal time; dots denote derivatives with respect to $t$ and primes denote derivatives with respect to $\eta$. \item $f_{\text{NL}}$ refers to $f_{\text{NL}}^\text{loc}$. \item $\rho_r$ refers to the total density in relativistic particles, which we calculate assuming that $n_\text{eff}=3.046$ and that neutrinos were relativistic through recombination. \item We use as $\zeta$ the same variable that \cite{Weinberg:2008zzc} calls $\mcl{R}$. Note that, in comparison with \cite{Shaw:2009nf}, $\zeta = -\zeta^\text{SL}$. \item Given that the Hubble expansion causes magnetic fields to decay as $B\propto a^{-2}$, we will often use a ``comoving'' magnetic field \begin{align} \tB \equiv B a^2\,, \end{align} and comoving energy density \begin{align} \tilde\rho \equiv \rho a^4\,, \end{align} so that $\tb$, $\tilde\rho$ are essentially normalized to their present values (e.g. $\tilde\rho_\gamma\approx\rho_{\gamma0}$). \item When we omit an explicit time dependence for a magnetic field, e.g. $\tB(\bk)$, we take it to be the field strength before any damping on relevant scales. \item We relate the two point function $\braket{\tb_i^*\tb_j}$ and $P_B$ via \begin{align}\label{eq:B-2_pt_fcn} \Braket{\tb_i^*(\bk_1)\tb_j(\bk_2)} &= (2\pi)^3 \frac{P_{ij}(\hat k_1)}{2} P_B(k_1) \delta(\bk_1 - \bk_2)\,; &&\text{where} & P_{ij}(\hat k) &=\delta_{ij}-\hat{k}_i\hat{k}_j\,, \end{align} we have neglected a helical component which we will not consider here. \item We refer to the damping scale at the start, finish of the $\mu$-distortion era as $k_D^i = 2.1\times10^{4}\text{ Mpc}^{-1}$, $k_D^f = 83 \text{ Mpc}^{-1}\,,$. We also define the power damping scale $\ck_D \equiv k_D/\sqrt{2}$. (See Sec. \ref{sec:damp-mgntc-flds} for more information). \item When performing calculations, we use values from Planck \cite{Planck-overview}. \item We assume that, besides the primordial magnetic field, the primordial perturbations are adiabatic. \end{itemize} | \label{sec:conclusion} In this work, we have investigated the sensitivity of the $\braket{\mu T}$ signal to a primordial $\bbzet$ correlation, analyzing the signal in upcoming experiments as well as considering the other possibly competing magnetic and density perturbation effects. We found that a CMBPol-like experiment can constrain $b_{\text{NL}} (\bmu/1\text{ nG})^2 \lesssim 100$, where $\bmu$ is the magnetic field on $\mu$-distortion scales. Thus, the $\braket{\mu T}$ correlation could meaningfully constrain primordial magnetic field correlations. We can also say with certainty that we will not have measurable post-magnetogenesis contributions to the $\bbzet$ correlator. The signal from the evolution of $\zeta$ due to magnetic fields, considered in Secs. \ref{sec:evln-of-zet-due-to-B}, \ref{sec:sign-from-evvn-of-zet} and previously in \cite{Miyamoto:2013oua}, is too small to affect a $b_{\text{NL}}$ measurement. We also saw in Sec. \ref{sec:evln-of-B-due-to-zet} that, due to the lack of superhorizon velocity flows in the early universe, we do not have a correlation induced by the evolution of $B$ due to initial perturbations. The small size of these effects is, perhaps, unsurprising because it is generally difficult to generate correlations between very different length scales. The signal from a squeezed-limit three-point function $f_{\text{NL}}$, the initial motivation for anisotropic $\mu$-distortion in \cite{Pajer&Zald:2012-New-window}, can potentially compete with a $b_{\text{NL}}$ signal, depending on the relative strengths of $f_{\text{NL}}$ and $b_{\text{NL}}$ and the size of the magnetic field and curvature perturbation on $\mu$-distortion scales. The shape of $C_l^{\mu T}$ is the same in both cases if both $f_{\text{NL}}$ and $b_{\text{NL}}$ are scale-invariant, so shape information cannot easily help us identify the source of a signal unless we have a motivated theoretical prior. In principle, one may be able to measure all of the $\mu$ monopole, $\braket{\mu T}$, and $\braket{\mu\mu}$ correlations; combining them, it might be possible to say more about the sources of a signal. We cannot neglect another important consideration: a magnetic field produced during inflation would, through its non-adiabatic pressure over the evolution of the universe, induce an $f_{\text{NL}}$. We discussed in Sec. \ref{sec:signal-from-fNL} that generically, for $\bmu\gtrsim 1$ nG, the induced $f_{\text{NL}}$ signal would dominate over the $b_{\text{NL}}$ signal from magnetic field damping. Since upcoming experiments will not be sensitive to such magnetic fields strengths unless $b_{\text{NL}}$ is large, the $f_{\text{NL}}$ signal induced from $b_{\text{NL}}$ would be more relevant for constraining magnetic fields in the near future. Note that this assumes the magnetic field is produced during inflation and then decays at the same rate as the rest of the universe. There are at least a few issues worth considering further. In particular, if one is interested in connecting the magnetic field strengths in the $\mu$-era to even earlier times, e.g. magnetogenesis, one should carefully explore the typically non-trivial evolution of magnetic fields in the early universe. It would also be interesting to consider if there are any other mechanisms that could produce a $\braket{\mu T}$ correlation besides a primordial $\bbzet$ correlation or density perturbation bispectrum. | 14 | 4 | 1404.5957 |
1404 | 1404.1372_arXiv.txt | {The massive infrared dark cloud G0.253+0.016 projected $\sim$45\,pc from the Galactic centre contains $\sim$10$^5$\,M$_{\odot}$ of dense gas whilst being mostly devoid of observed star-formation tracers.} {Our goals are therefore to scrutinise the physical properties, dynamics and structure of this cloud with reference to its star-forming potential.} {We have carried out a concerted SMA and IRAM 30m study of this enigmatic cloud in dust continuum, CO isotopologues, several shock tracing molecules, as well as H$_2$CO to trace the gas temperature. In addition, we include ancillary far-IR and sub-mm \textit{Herschel} and SCUBA data in our analysis.} {We detect and characterise a total of 36 dust cores within G0.253+0.016 at 1.3\,mm and 1.37\,mm, with masses between 25 and approximately 250\,M$_{\odot}$, and find that the kinetic temperature of the gas traced by H$_2$CO ratios is $>$320\,K on size-scales of $\sim$0.15\,pc. Analysis of the position-velocity diagrams of our observed lines shows broad linewidths and strong shock emission in the south of the cloud, indicating that G0.253+0.016 is colliding with another cloud at v$_{_{\rm LSR}}\sim 70$\,km\,s$^{-1}$. We confirm via an analysis of the observed dynamics in the Central Molecular Zone that it is an elongated structure, orientated with Sgr\,B2 closer to the Sun than Sgr\,A*, however our results suggest that the actual geometry may be more complex than an elliptical ring. We find that the column density Probability Distribution Function (PDF) of G0.253+0.016 derived from SMA and SCUBA dust continuum emission is log-normal with no discernible power-law tail, consistent with little star formation, and that its width can be explained in the framework of theory predicting the density structure of clouds created by supersonic, magnetised turbulence. We also present the $\Delta$-variance spectrum of this region, a proxy for the density power spectrum of the cloud, and show it is consistent with that expected for clouds with no current star formation. Finally, we show that even after determining a scaled column density threshold for star formation by incorporating the effects of the increased turbulence in the cloud, we would still expect ten stars with masses $>$15\,M$_{\odot}$ to form in G0.253+0.016. If these cannot be accounted for by new radio continuum observations, then further physical aspects may be important, such as the background column density level, which would turn an absolute column density threshold for star formation into a critical over-density.} {We conclude that G0.253+0.016 contains high-temperatures and wide-spread shocks, displaying evidence of interaction with a nearby cloud which we identify at v$_{_{\rm LSR}}\sim 70$\,km\,s$^{-1}$. Our analysis of the structure of the cloud can be well-explained by theory of magnetised turbulence, and is consistent with little or no current star formation. Using G0.253+0.016 as a test-bed of the conditions required for star formation in a different physical environment to that of nearby clouds, we also conclude that there is not one column density threshold for star formation, but instead this value is dependant on the local physical conditions.} | } Determining how massive clusters ($10^{3} - 10^{5}$\,M$_{\odot}$) form has a profound effect on how we interpret observations of star formation in external galaxies. As the majority of stars form in clusters \citep{lada03,de-wit05}, and because massive clusters yield -- either via statistics or by virtue of their physical conditions -- the most massive stars, these clusters are the engines which produce the objects that dominate the luminosity of galaxies. Therefore, uncovering how and where these massive clusters can form is of central importance in understanding how galaxies evolve, and may provide hints as to how cluster formation proceeds at all masses. To understand the conditions that lead to cluster formation, it is necessary to observe the structure of the gas and dust, as well as kinematics, of a cluster forming cloud before star formation processes begin blurring the initial conditions. Observing the formation of massive clusters within our own Galaxy has the obvious advantage of resolving clouds that typically fall within a single resolution element for observations of distant galaxies. One of the most exceptional candidates for a massive cluster progenitor is the cloud G0.253+0.016 near the Galactic centre (e.g., \citealt{lis94,lis98,longmore12}). The global dust properties of G0.253+0.016 (also known as M0.25+0.01) have been shown by previous authors \citep[e.g.][]{lis94,lis98,lis01,longmore12, immer12} to be cold ($\sim$18-27\,K), dense (n$\sim$7.3\,$\times$10$^{4}$ - 6\,$\times$10$^{5}$ cm$^{-3}$) and massive (M=1.3 - 7 $\times10^{5}$\,M$_{\odot}$). However, minimal evidence has been found for ongoing star formation. The H$_2$O maser and 8.4\,GHz radio continuum observations of \citet{lis94} uncovered a single H$_2$O maser near the 350\,$\mu$m continuum peak position, but no coincident radio emission. \textit{Spitzer} and \textit{Herschel} observations have also confirmed that no protostars are seen in the cloud in the mid- or far-IR up to 70\,$\mu$m \citep[e.g.,][]{longmore12}. \citet{rodriguez13} presented Very Large Array (VLA) 1.3 and 5.6\,cm radio continuum observations of G0.253+0.016 which detected three compact radio sources towards the eastern edge of the cloud, which have thermal spectral indices and correspond to B0.5 stars. These could be signposts of massive star formation in G0.253+0.016; however, as will be discussed below, they do not correspond to dense cores of gas traced by mm emission. It is interesting to note that the similarly massive cloud Sgr B2, which is not far away from G0.253+0.016, actively forms stars and is one of the most prominent sites of star formation in the Galaxy. Therefore, how can such a massive and dense molecular cloud as G0.253+0.016 currently not be forming stars? A possible explanation for the high star-formation activity in Sgr B2 is that one of the dust lanes associated with the Galactic bar intersects with the $x_{2}$ orbits of the gas in the Central Molecular Zone (CMZ) at the position of Sgr B2, and thus we would expect an elevated level of star formation at this position. Yet this does not straightforwardly explain the comparatively lesser degree of star formation towards the other point of intersection with the dust lanes of the bar on the opposite side of the CMZ, Sgr C \citep[e.g.,][]{kendrew13}. One possible explanation could involve an asymmetry in the material falling inwards along the two dust lanes of the bar, or in the gas orbiting the CMZ, providing less material for collision at the position of Sgr C. Alternatively, the star formation in Sgr B2 could be enhanced by its recent passage close to the Galaxy's central black hole, Sgr\,A* \citep{longmore13}. Of course, the observed differences between Sgr B2 and Sgr C may be due to a combination of these effects. Recently, G0.253+0.016 has also been studied using MALT90 and APEX observations \citep{rathborne14}, and ALMA SO observations have shown evidence for a cloud-cloud collision with G0.253+0.016 \citep{higuchi14}. In fact, such cloud-cloud collisions may provide a way to collect enough dense gas to produce massive clusters \citep[e.g.,][]{fukui14}. With reference to these possible scenarios, in this work we aim to determine the current state of star formation in G0.253+0.016, as well as its star-forming fate (i.e., whether it will form a star cluster), by investigating the structure of the cloud, its internal dynamics, its interaction with the CMZ environment and its local stability to collapse. Section \ref{obs} outlines our 1.3\,mm Submillimeter Array (SMA) and Institut de Radioastronomie Millim\'{e}trique (IRAM) 30m observations, as well as ancillary far-IR and sub-mm \textit{Herschel} and Submillimeter Common-User Bolometer Array \citep[SCUBA;][]{holland99} data. Section \ref{result} presents our observational results from the continuum and line observations, including the column density Probability Distribution Function (PDF) of G0.253+0.016, and determination of temperatures from H$_2$CO. Section \ref{discussion} presents our discussion, which covers the topics of the internal dynamics of G0.253+0.016, the interaction of G0.253+0.016 with its environment, as well as its current star-formation activity and potential. Our conclusions are given in Section~\ref{conclusions}. | \label{conclusions} To scrutinise the dynamics and structure, as well as determine the star-forming potential, of the massive Galactic centre cloud G0.253+0.016, we have carried out a concerted SMA and IRAM 30m study of this enigmatic cloud in dust continuum, CO isotopologues, several shock tracing molecules such as CH$_3$OH and SiO, as well as H$_2$CO to trace the gas temperature. In our study, we have also included ancillary far-IR and sub-mm \textit{Herschel} and SCUBA data to further the interpretation of the cloud structure. Our main results are as follows: \begin{enumerate} \item We have characterised the 36 cores detected in G0.253+0.016 with the SMA in 1.3\,mm and 1.37\,mm continuum, with masses between 25 and approximately 250\,M$_{\odot}$, comparing these to recent 1.3\,cm VLA observations (Mills et al., in preparation) which suggest that several cores are associated with free-free ionised gas continuum and thus may be tracing the first signs of massive star formation in this cloud. \item By modelling H$_2$CO line ratios, we find that the kinetic temperature of the gas is extremely large ($>$320\,K) on the size-scales traced by the SMA beam ($\sim$4.3$\times$2.9$''$ or 0.18$\times$0.12\,pc). These temperatures are much hotter than those found for that of the dust, which we find to reach below 20\,K in the innermost regions of the cloud, in agreement with previous results. \item We observe widespread shock emission over G0.253+0.016, which is strongest in the southern regions of the cloud. Further, by comparing position-velocity diagrams of the large-scale $^{13}$CO emission to that of the shock tracing molecules, we find that G0.253+0.016 intersects in position and velocity with another cloud which peaks at 70\,km\,s$^{-1}$. The shock tracers are brightest and display large velocity gradients close to this interaction zone, indicating cloud-cloud collision. Hence, we have found the first evidence of the specific cloud colliding with G0.253+0.016. \item By investigating the dynamics of the entire Galactic centre region, we find that the HNC Galactic longitude-velocity diagram of the CMZ is consistent with the CMZ being orientated with Sgr\,B2 on the near-side. We also determine that if this cloud is in fact colliding with another cloud at 70\,km\,s$^{-1}$, a more complex geometry for the CMZ is required than a simple elliptical ring structure. \item We determined the column density PDF of G0.253+0.016 derived from SMA and SCUBA dust continuum emission is log-normal with no discernible power-law tail, consistent with little star formation. We also find that the width of the column density PDF is narrow but that, given the level of turbulence in the Galactic centre and the enhanced magnetic field in this region, it is consistent with the expected column density PDF created by supersonic, magnetised turbulence. \item We also investigate the $\Delta$-variance spectrum of this region and show it is consistent with that expected for clouds with 0\% SFE, supporting the fact we see no evidence for star formation from the column density PDF. \item We show via a simple argument using the virial mass that the star formation column density threshold for G0.253+0.016 should be increased due to the increased turbulence in the CMZ compared to clouds in the Galactic disk, yet incorporating turbulence might still not account for the lack of massive star formation observed. In addition to the level of turbulence, the background or average density may also play a role in the determination of the local column density threshold, making it instead a critical over-density. Thus we confirm that there is no one column density threshold for star formation, but this is dependant on local cloud conditions, and suggest that the Kennicutt-Schmidt relation for external galaxies is observed simply because galactic disk clouds have similar average densities and levels of turbulence, and dominate over clouds in the galactic nucleus such as G0.253+0.016. \end{enumerate} \begin{figure} \includegraphics[width=9.cm]{figures/fig18.pdf} \caption{The $\Delta$-variance spectrum of G0.253+0.016 as a function of the fractional length scale $\ell /L$, calculated from the column densities derived from the SCUBA-only image (grey line), and from the combined SMA plus SCUBA image (black line). The displayed errors are calculated via the method outlined in \citet{bensch01}. The red line shows the best fit to the combined SMA plus SCUBA $\Delta$-variance between 0.08 and 0.2, giving a slope $1.907\pm0.037$ corresponding to $\alpha=-1.907\pm0.037$. \label{deltavar}} \end{figure} | 14 | 4 | 1404.1372 |
1404 | 1404.3970_arXiv.txt | We have obtained low and medium resolution spectra of 9 brown dwarf candidate members of Coma Berenices and the Hyades using SpEX on the NASA InfaRed Telescope Facility and LIRIS on the William Herschel Telescope. We conclude that 7 of these objects are indeed late M or early L dwarfs, and that two are likely members of Coma Berenices, and four of the Hyades. Two objects, cbd40 and Hy3 are suggested to be a field L dwarfs, although there is also a possibility that Hy3 is an unresolved binary belonging to the cluster. These objects have masses between 71 and 53 M$_{\rm Jup}$, close to the hydrogen burning boundary for these clusters, however only an optical detection of Lithium can confirm if they are truly substellar. | The Hyades (Melotte 25; RA= 04 26.9, Dec= +15 52) and Coma Berenices (Melotte 111; RA = 12 23 00, Dec = +26 00 00, J2000.0) are the closest open star clusters to the Sun at distances of 46.45$\pm$0.5 and 86.7$\pm$0.9 pc respectively \citep{vanleuwen09}. Both these clusters are relatively mature at 625$\pm$50 Myr \citep{perryman98} and $\sim$500 Myr, and have low reddening values of E(B-V)$<$1.0 and 3.2 mmag respectively \citep{taylor06}, but this is where the similarities end. The Hyades has been well studied (e.g. \citealt{reid92,gizis99,dobbie02}), and contains many members ($\sim$400), whereas Coma Ber has been the subject of relatively few detailed studies and membership is less certain. One of the reasons Coma Ber is often neglected is that it covers a relatively large region on the sky ($\sim$100 deg$^{2}$), while being sparsely populated. A low proper motion (-11.5, -9.5 mas yr$^{-1}$) also makes kinematic surveys with a baseline of less than 10 years difficult. We performed a wide area search for new stellar members in \citet{casewell06}, identifying 60 candidates. A more recent kinematic and photometric survey by \citet{kraus07} discovered 149 candidate members, of which 98 have a membership probability of $>$80 per cent, and determined that this survey was complete to 90 per cent between the spectral types of F5 and M6. A search for substellar members was also performed by \citet{casewell05} who identified 13 new brown dwarf candidates using optical and near-IR photometry. A similar optical photometric survey was performed by \citet{melnikov12} who surveyed 22.5 deg $^{2}$ and combined $RI$ photometry with 2MASS and UKIDSS. They discovered 12 new low mass members down to a spectral type of M6-8. \citet{terrien14} also performed a photometric and proper motion survey of the cluster, discovering 8 new stellar members, and confirming the membership of 6 M dwarfs discovered by \citet{kraus07} by measuring their radial velocity. \citet{mermilliod08} also performed a radial velocity study of 69 solar type stars, 46 of which were candidate cluster members, to search for close, low mass companions. Of these 46, only 8 stars appear to be cluster members, and 6 additonal members were determined to be spectroscopic binaries, suggesting a spectroscopic binary fraction of 22\%. Coma Ber is only estimated to have 145$\pm$15 stars earlier than M6 \citep{kraus07} and a total mass of 112$\pm$16 M$_{\odot}$ compared to a total mass of 300-460 M$_{\odot}$ for the Hyades \citep{reid92}. These mature clusters are expected to have undergone some form of mass segregation as part of their dynamical evolution, with some sources suggesting that as many as 50 per cent of the low mass members being lost with time \citep{delafuente00}. In their large area study of Coma Ber \citet{kraus07} suggest that some mass loss has occurred, but over a lower mass range than their survey covered. Despite there being clear evidence of mass segregation in Praesepe, a similarly aged cluster, recent surveys have identified many new candidate brown dwarfs \citep{baker10, boudreault12}, one of which has recently been spectroscopically confirmed \citep{boudreault13}. There are also two known T dwarf members of the Hyades \citep{bouvier08}, and \citet{goldman13} reports the discovery of 43 new cluster members, many of which are low mass, but are not in the substellar regime. They also suggest that one previously known L0 dwarf, 2MASSIJ02330155+270406 \citep{cruz07} is also a cluster member. These open clusters provide excellent laboratories for studying brown dwarfs, mainly due to their coeval nature and known distances, ages and metallicities. Using brown dwarfs found in open clusters as benchmark objects is not new, as evidenced by the many field objects that are discovered, and latter associated with moving groups (e.g. \citealt{jameson08, casewell08}). Recent work on the field population has separated brown dwarfs into gravity categories, using the suffixes $\alpha$, $\beta$, $\gamma$ and $\delta$ \citep{cruz09}. Where $\alpha$ is used to denote a field or "normal" gravity object, $\beta$ is used for an intermediate gravity object (~100 Myr), $\gamma$ a low gravity object (10-30 Myr) and $\delta$ a very low gravity object (~1 Myr). At the age of the Hyades and Coma Ber however, the gravity is high enough to be very similar to, but slightly lower than that of field dwarfs \citep{chabrier00}. This does, however, mean that any object with a lower than average gravity can be discounted as a potential cluster member. We recently performed proper motion and photometric searches of both the Hyades and Coma Ber clusters to discover new substellar candidate members \citep{hogan08,casewell05}. These surveys identified 12 new Hyades L dwarf candidates (using the moving group method and near-IR photometry) and 13 new brown dwarf candidates (using optical and near-IR photometry only) in Coma Ber. These objects, if bona fide cluster members can be used as benchmarks as their age and metallicity are known, and they provide a sample of brown dwarfs with a near, but lower than field gravity. In this paper we present near-infrared spectra of 9 of these candidates. | Using near-IR spectra from SpeX on IRTF and LIRIS on the WHT we have obtained spectra of 9 brown dwarf candidate members of the Hyades and Coma Ber. We have rejected cbd36 and Hy2 as brown dwarfs from their spectra. We also reject cbd40 as a member of Coma Ber, and suggest that Hy3 may be a field object, or an unresolved binary that is a cluster member. The remaining objects have spectral types ranging from M9 to L2 and masses between 71 and 53 M$_{\rm Jup}$. Using EWs and the indices defined by \citet{allers13} we have determined that none of these objects have low gravity, which would indicate that they are younger than the clusters they have been identified with, thus supporting their cluster membership. These are the first spectroscopically confirmed L dwarfs in the Hyades cluster and Coma Ber, however, optical spectra containing a lithium detection is required to determine if these objects have a truly substellar nature. | 14 | 4 | 1404.3970 |
1404 | 1404.3210.txt | The recurrent nova T~Pyx underwent its sixth historical outburst in 2011, and became the subject of an intensive multi-wavelength observational campaign. We analyze data from the \emph{Swift} and \emph{Suzaku} satellites to produce a detailed X-ray light curve augmented by epochs of spectral information. X-ray observations yield mostly non-detections in the first four months of outburst, but both a super-soft and hard X-ray component rise rapidly after Day 115. The super-soft X-ray component, attributable to the photosphere of the nuclear-burning white dwarf, is relatively cool ($\sim$45 eV) and implies that the white dwarf in T~Pyx is significantly below the Chandrasekhar mass ($\sim$1 M$_{\odot}$). The late turn-on time of the super-soft component yields a large nova ejecta mass ($\gtrsim10^{-5}$ M$_{\odot}$), consistent with estimates at other wavelengths. The hard X-ray component is well fit by a $\sim$1 keV thermal plasma, and is attributed to shocks internal to the 2011 nova ejecta. The presence of a strong oxygen line in this thermal plasma on Day 194 requires a significantly super-solar abundance of oxygen and implies that the ejecta are polluted by white dwarf material. The X-ray light curve can be explained by a dual-phase ejection, with a significant delay between the first and second ejection phases, and the second ejection finally released two months after outburst. A delayed ejection is consistent with optical and radio observations of T~Pyx, but the physical mechanism producing such a delay remains a mystery. | The five thermonuclear explosions of T~Pyxidis observed in 1890, 1902, 1920, 1944 and 1966 earned the system its place as the prototypical recurrent nova, but have also highlighted our poor understanding of many aspects of binary evolution and nova theory. The community has waited anxiously for the next outburst of T~Pyx in order to study this peculiar system with modern multi-wavelength capabilities, and T~Pyx finally obliged by entering its sixth recorded outburst in April 2011. High-quality panchromatic observations are now revealing a host of new surprises for this system. A nova is a transient event marking a thermonuclear runaway on the surface of an accreting white dwarf. The white dwarf accretes hydrogen-rich material from a companion star, and this accreted material settles down into a thin degenerate layer on the surface of the white dwarf. The pressure and temperature in this layer increase until explosive nuclear burning begins, and the bulk of the accreted envelope is expelled from the white dwarf at hundreds to thousands of km s$^{-1}$. Novae are expected to recur on an accreting white dwarf with a timescale primarily determined by the white dwarf mass and accretion rate \citep[e.g.,][]{Yaron_etal05, Wolf13}. Predicted recurrence timescales vary widely ($\sim$1--10$^{8}$ years; \citealt{Yaron_etal05}), and those novae repeating on historical timescales have been dubbed ``recurrent" novae. Theoretically, we expect recurrent novae to occur in binaries where massive white dwarfs accrete at high rates, because more massive white dwarfs have higher surface gravities, meaning that the critical conditions are reached for smaller accreted envelopes, and higher accretion rate systems accrue this trigger mass in less time. Even before 2011, T~Pyx flew in the face of our expectations for recurrent novae. The evolution of its optical light curve is slow, showing a several months-long plateau around maximum light and a relatively slow decline from this maximum \citep{Schaefer10b}. T~Pyx has a short orbital period (1.83 hours; \citealt{Uthas10}), solidly below the cataclysmic-variable (CV) period gap. According to the theory of CV evolution, such short-period systems should have, on average, very low accretion rates \citep[e.g.,][]{Knigge11a}, but observations in quiescence---and the short nova recurrence time---imply that T~Pyx has an accretion rate orders of magnitude higher than these expectations \citep{Gilmozzi_Selvelli07, Selvelli08}. In addition, measured binary parameters imply that the white dwarf in T~Pyx may be significantly less massive than the Chandrasekhar mass \citep{Uthas10}, in contrast with common assumptions for recurrent novae. The high accretion rate in T~Pyx does not appear to be sustainable, as it is exceeds expectations by several orders of magnitude for mass transfer rates driven by gravitational radiation (the commonly-accepted mass transfer mechanism at such short orbital periods). \citet{Knigge00} and \citet{Schaefer10a} have hypothesized that mass transfer in T~Pyx is in a short-term elevated state, perhaps incited by a powerful nova outburst which occurred during the 1800s (before regular records were kept on T~Pyx). Before this postulated event, T~Pyx may have been a typical CV below the period gap, with a very low accretion rate and long intervals between novae. However, after the hypothesized explosion, the hot white dwarf irradiated the companion star and induced an unusually high mass-transfer rate. Perhaps this irradiation power is slowly dwindling and the accretion rate is gradually declining, explaining the increasing intervals between nova events observed for T~Pyx throughout the last century. Observational tests of this hypothesis have reached divergent conclusions as to whether there is evidence for a secular decline in T~Pyx's accretion rate \citep[e.g.,][]{Schaefer13, Godon14}. \begin{figure*}[t] \begin{center} \includegraphics[height=6.8in]{tpyx_rxo_lcurve_lessfreqs.ps} \caption{Overview of the 2011 outburst of T~Pyx at optical, X-ray, and radio wavelengths. {\it Top:} V-band optical data from the AAVSO. {\it Middle:} 0.3--10 keV X-ray light curve obtained with {\it Swift}/XRT. {\it Bottom: } VLA radio light curve at 2.5, 7, and 37 GHz (Paper I). } \label{rxo} \end{center} \end{figure*} Regardless of T~Pyx's history, it is clear that T~Pyx provides an opportunity to test an unusual corner of nova parameter space. Compared with other novae, we have a thorough understanding of the binary system's parameters and the accretion rate \citep{Selvelli08, Uthas10}. With multi-wavelength data collected from the 2011 outburst, we can measure key properties of the nova event, like ejected mass, and compare them with predictions from nova models. With this goal in mind, campaigns have been carried out across the entire electromagnetic spectrum, providing an exquisitely detailed picture of the 2011 nova outburst of T~Pyx \citep{Chesneau11, Kuulkers11b, Kuulkers11a, Shore11, Shore13, Evans12, Imamura_Tanabe12, Nelson12t, Ederoclite13, Williams13, Schaefer13, Patterson13, Sokoloski13, Tofflemire13, DeGennaroAquino14, Godon14, Surina14}. At every wavelength studied so far, the 2011 outburst of T~Pyx shows surprising features when compared to expectations for recurrent novae. As in previous outbursts, the optical light curve shows a sort of plateau for three months (Figure \ref{rxo}), implying that the optical photosphere is roughly constant in size for $\sim$90 days after thermonuclear runaway \citep{Shore13}. Optical spectroscopy shows that T~Pyx is an unusual ``hyper-hybrid" nova, switching from He/N class to \ion{Fe}{2} class around Day 10, and then back to He/N on Day 65 \citep{Williams12, Williams13, Ederoclite13, Surina14}. The radio light curve and optical measurements of the change in the binary period after the nova event imply a large ejected mass ($\sim$10$^{-4}-10^{-5}$ M$_{\odot}$; \citealt{Nelson12t} [henceforth Paper I], \citealt{Patterson13}), rather than the $\sim$10$^{-6}-10^{-7}$ M$_{\odot}$ expected for recurrent novae. Radio light curves also show a late and steep rise (Figure \ref{rxo}; Paper I), implying that either the ejecta in T~Pyx were very cold ($<$200 K) during the first $\sim$50 days of the outburst, or the bulk of the nova ejecta stalled at an $\sim$AU-scale radius until it was finally expelled $\sim$50 days after the thermonuclear runaway (such a delay might also explain the long plateau in the optical light curve; \citealt{Shore13}). In this work, we focus on X-ray observations of the 2011 outburst of T~Pyx obtained with {\it Swift} and {\it Suzaku}, and compare our results with inferences from other wavelengths. X-ray emission from novae can be split into two broad classes, which may, but need not, exist contemporaneously \citep{Krautter08}. The first class is super-soft X-ray emission, characterized by effective temperatures between 10$^{5}$ and 10$^{6}$ K, and luminosities in the range 10$^{36}$--10$^{38}$ erg s$^{-1}$. High resolution spectra obtained with the grating instruments onboard {\it Chandra} and {\it XMM-Newton} have confirmed that this emission originates near the white dwarf photosphere \citep{Nelson08, Rauch10, Ness11, Orio12}. In some novae (e.g., V2491 Cyg, RS Oph) the soft X-ray flux is continuum emission that most likely originates at the white dwarf photosphere. In other cases, the soft X-rays are associated with strong lines of H and He-like carbon, nitrogen and oxygen that likely indicate scattered photospheric emission \citep[see ][and references therein]{Ness13}. Super-soft X-ray emission only becomes visible at later stages of the nova outburst, once the ejecta have become optically thin to the radiation from the hot, still burning white-dwarf surface layers; therefore, the emergence time of the super-soft source can be used as a diagnostic for ejecta mass \citep{Henze11, Schwarz11}. Observations of novae in the $\sim$1--10 keV energy range also reveal harder X-ray emission on timescales of days to years after outburst \citep[e.g.,][]{Mukai08}. During the novae in RS~Oph and V407~Cyg, both of which have red giant secondaries, hard X-ray emission was detected at early times and attributed to the interaction of the nova ejecta with the dense wind of the companion \citep{Sokoloski06, Nelson12}. In systems with less evolved donors (and hence lower density circumbinary environments), internal shocks within the ejecta have been proposed as the origin for hard X-ray emission \citep[see][and references therein]{Mukai08}. In these cases, hard X-ray emission can tell us about the structure of the ejecta as the nova outburst progresses \citep{O'Brien94}. In the last three years, novae have been identified as a new class of GeV gamma-ray transient by {\it Fermi}/LAT, indicating that the shock interactions in novae are capable of accelerating particles to relativistic speeds (\citealt{Abdo10, Hill13}; see also \citealt{Tatischeff_Hernanz07}). Observations of novae in the 1--10 keV range are required to characterize these shocks and fully understand the gamma-ray production mechanism. We note that T~Pyx was \emph{not} detected with \emph{Fermi} (C.~Cheung 2013, private communication), making it a useful comparison case for the study of why some novae produce detectable gamma-rays while others do not. In this paper, we discuss both modes of X-ray emission during the 2011 outburst of T~Pyx. In Section 2 we discuss the X-ray observations and data reduction; \emph{Swift} monitoring reveals the X-ray evolution at high cadence, while our single epoch of \emph{Suzaku} spectroscopy provides high signal-to-noise on Day 194. In Section 3, we describe the \emph{Swift} X-ray light curve, and Section 4 presents our spectral analysis of the \emph{Swift} and \emph{Suzaku} data. Section 5 analyzes the observed hard X-ray component and concludes that it is likely produced by a shock within the ejecta, rather than interaction between the nova and pre-existing circumbinary material. Section 6 presents a super-soft component with a relatively cool temperature and late turn-on time (compared to other recurrent novae). In Section 7, we discuss how these X-ray results align with optical and radio observations, and suggest that all three wavelength regimes support a second, massive, delayed ejection in T~Pyx. We conclude in Section~8. Throughout the paper, we take 2011 April 14 (MJD = 55665) as $t_{0}$ or Day 0, the beginning of optical rise and the start of the outburst \citep{Waagan11, Schaefer13}. We also assume a distance to T~Pyx of $4.8\pm0.5$ kpc \citep{Sokoloski13}. | In our analysis of \emph{Swift} and \emph{Suzaku} data covering the 2011 outburst of the recurrent nova T~Pyx, we detect two distinct components of X-ray emission: a super-soft component associated with the photosphere of the nuclear-burning white dwarf, and a hard component associated with shocked thermal gas. The super-soft X-ray component becomes detectable between Days 123--143, implying an ejection which is surprisingly massive for a recurrent nova ($\gtrsim 10^{-5}$ M$_{\odot}$), but consistent with other recent constraints for T~Pyx \citep[Paper I]{Selvelli08, Patterson13}. In addition, the temperature of the super-soft source is relatively cool (30--50 eV), implying that the white dwarf in T~Pyx is significantly below the Chandrasekhar mass. T~Pyx therefore inhabits a poorly-explored corner of nova parameter space, where its unusually high accretion rate, rather than an unusually massive white dwarf, drives it to have a short nova recurrence time. A hard X-ray component is also detected in all epochs with sufficient signal-to-noise for spectral analysis (Days 142--206). Hard X-rays are relatively common in novae, and are postulated to originate in shocks internal to the nova ejecta \citep{O'Brien94, Krautter08}. The hard X-rays in T~Pyx, however, are accompanied by an unusually strong \ion{O}{8} Ly$\alpha$ emission line, which can only be fit if the thermal plasma has a highly super-solar abundance of oxygen. Such high oxygen abundances in the nova ejecta require the dredge-up of significant amounts of white dwarf materials during the nova event (contrary to some models of recurrent novae; e.g., \citealt{Starrfield12}). Similar conclusions can be reached from analysis of \emph{Chandra} LETG spectra, which find a factor of $\sim$15 overabundance of nitrogen in the ejecta, relative to solar (consistent with enrichment by CNO-processed white dwarf material; \citealt{Tofflemire13}). In addition, we note the appearance of a faint, hard X-ray component at early times with an unknown origin. This component was only detectable by \emph{Swift}/XRT in time-averaged spectra spanning Days 14--20, and its low signal-to-noise precludes detailed spectral analysis. However, we note a single coincident radio detection with the VLA at 33 GHz on Day 17 (Paper I). The nature of this component remains a mystery, although it may be linked to the first low-mass ejection from T~Pyx on Day 0 (Figure \ref{cartoon}). As measured from optical emission line profiles, the expansion velocity of ejecta in T~Pyx increases by 50\% during the first two months of outburst (Paper I, \citealt{Surina14}); we find that this variation, and subsequent interaction within the ejecta, can naturally explain the temperature and light curve of the hard X-ray component. In Section \ref{multi}, we present a cartoon picture that can self-consistently explain the soft and hard X-ray evolution, the radio light curve, the optical spectral evolution, and the optical light curve. A shell of material is expelled on Day 0 at 1,900 km s$^{-1}$, and a second episode of mass ejection is released on Day $\sim$65 at 3,000 km s$^{-1}$ (Figure \ref{cartoon}). While the first ejection produces clearer signatures at optical wavelengths, the second ejection accounts for the bulk of the expelled mass (as implied by the small absorbing column in the hard X-ray component and the bright but delayed maximum in the radio light curve). We do not yet understand the physical mechanism that leads to the bulk of the nova mass stalling out and lingering around the binary for two months before finally being expelled, but T~Pyx now joins a significant and growing sample of novae which show evidence for complex, multi-phase mass ejection (see \citealt{Lynch08, Krauss11, Williams12} for other examples). | 14 | 4 | 1404.3210 |
1404 | 1404.1833_arXiv.txt | We sketch the average behaviour of the temperatures and densities of the main components of the $\Lambda$CDM universe after inflation. It is modelled as a perfect fluid with dark energy associated with the macroscopic effect of conformal variations of the metric. The main events of the thermal evolution are studied, such as the effect of particle annihilations and decoupling, and the transitions between the eras dominated by different entities. Estimates of the average present epoch temperature of baryonic matter and dark matter composed of neutralinos are given. We study the eventual presence of a sterile neutrino component and find that the sterile neutrino density at the epoch of primordial nucleosynthesis is in agreement with expectations when their evolution starts, at the end of inflation, in temperature equilibrium with the rest of the universe. | The aim of this paper is to study the evolution of the scale factor and the global behaviour of temperatures and densities of the components of the universe after the end of the epoch of inflation and how these quantities are influenced by the property of the neutrinos being Majorana or Dirac particles and by the choice of dark matter candidate masses. In this description of the average evolution neither the details of the primordial nucleosynthesis nor the impact of structure formation on the matter temperature are taken into consideration. We assume dark matter to be of only one type and presume that the chemical potentials are negligible in the regimes we are interested in.\\ To track the thermal history of the universe, let us consider the standard cosmological picture based on the observational evidence that the universe is highly homogeneous, isotropic and flat, and composed of approximately 72\% dark energy, 23\% dark matter and 5\% baryonic matter \cite{WMAP7}.\\ Near the end of the twentieth century, astronomical observations of redshifts of supernovae \cite{perlmutter, riess} have led to the conclusion that the universe is expanding at an accelerated rate. This behaviour indicates the presence of an entity with negative pressure, the dark energy, and the concept of a cosmological constant in Eintein's equations has thus been revived. It acts as a repulsive gravity responsible for the accelerated expansion of the universe.\\ One way to account for the origin and nature of the cosmological constant is by considering quantum fluctuations of the metric \cite{narlikar, Blin}. At the Planck scale, gravity and quantum mechanics can not be seen as independent. Space-time is expected to be coarse grained, requiring a quantum metric tensor.\\ The simplest way of generalizing the metric to include quantum effects is through a scalar field $\varphi$ describing conformal variations of the metric around its classical value. The advantage of conformal variations is that causality is obeyed, since the light cone structure remains intact. These fluctuations can be written as follows \cite{narlikar, Blin} \begin{equation} g_{\mu \nu} = (1+\varphi)^2\bar{g}_{\mu \nu}=\Phi^2\bar{g}_{\mu \nu}, \end{equation} where $\bar{g}_{\mu \nu}$ represents the usual classical metric tensor about which the fluctuations occur. The fluctuation average is required to be $<\varphi>=<\varphi_{, \mu}>=0$, centring the generalized metric around its classical value and yielding no drift of $\varphi$ in space-time.\\ The presence of fluctuations changes the Ricci tensor $R_{\mu \nu}$ and curvature scalar $R$, and when deducing Einstein's equations from variation of the Einstein-Hilbert action with respect to the classical metric {\sl and} the fluctuation field, one gets (units $\hbar=c=1$)\cite{Blin} \begin{equation} \label{eq:maisuma} \bar{R}_{\lambda \mu} - \frac{1}{2}\bar{g}^{\lambda \mu}\bar{R} - \bar{g}_{\lambda \mu}\Lambda= 8 \pi G T_{\lambda \mu}, \end{equation} with \begin{equation} \label{eq:Lambda} \Lambda = -\frac{1}{4}\left(\bar{R} + 8\pi G T_{\lambda \mu}\bar{g}^{\lambda \mu}\right), \end{equation} where the quantities with a bar refer to the classical expressions without fluctuations. Equation (\ref{eq:maisuma}) is Einstein's equation with a cosmological constant (\ref{eq:Lambda}) that stems from the fluctuations. It is possible to show \cite{Blin} that the magnitude of $\Lambda$ is consistent with the observational values attributed to dark energy. Thus, the accelerated expansion that dark energy accounts for in standard cosmology can be seen as a consequence of considering quantum fluctuations present at Planck scale. \\ For the universe as a perfect fluid, the stress-energy tensor has the form \begin{equation} \label{eq:Tmunu} T^{\mu \nu} = \left(\rho + P\right)u^\mu u^\nu - Pg^{\mu \nu}, \end{equation} where $\rho$ represents the mass-energy density and $P$ represents the pressure of the fluid. We can use it in equation (\ref{eq:Lambda}) to ascertain that the cosmological constant derived this way is indeed constant. Taking the time derivative of equation (\ref{eq:Lambda}), we confirm \cite{tese} that $\dot{\Lambda} =0$. \\ We will use the Robertson-Walker metric, appropriate for a homogeneous and isotropic universe, as observations seem to indicate ours is, \begin{equation} ds^2=dt^2-Q(t)^2\left[\frac{dr^2}{1-kr^2}+r^2(d\theta^2+\sin^2\theta d\phi^2)\right], \end{equation} where $Q(t)$ represents the scale factor and $k$ the curvature of the space.\\ The standard model for cosmology stipulates that the universe originated from a hot dense state, that expanded and cooled down with time. This Big Bang model can be complemented by inflationary theories, which propose a brief period of exponential growth for early times. We can study what happened after inflation, by using Friedmann's equation \begin{equation} \label{eq:00} 3\frac{\dot{Q}^2 + k}{Q^2} - \Lambda = 8\pi G \rho. \end{equation} It is customary to divide the density by the critical density ($H_0$ is the Hubble constant) \begin{equation} \label{eq:rho_c} \rho_c=\frac{3H_0^2}{8 \pi G}, \end{equation} and work with the density parameter $\Omega=\rho/\rho_c$. We can define different density parameters for the different entities, namely the density parameter for matter, $\Omega_{m}$ (where we can further distinguish between the density of baryonic matter $\Omega_{b}$ and cold dark matter $\Omega_{c}$) and for radiation, $\Omega_{r}$. Similarly, the cosmological constant and curvature can be expressed by the density parameters $\Omega_{\Lambda}=\Lambda/3H_0^2$ and $\Omega_{k}= -k/(Q(t)^2H_0^2)$.\\ It will also be important to consider the equation that describes the energy-mass conservation \begin{equation} \label{eq:rhoponto} \dot{\rho}_i = -3\frac{\dot{Q}}{Q}(\rho_i + P_i), \end{equation} which can be written independently for each component $i$, as long as there is no significant conversion between them. \\ We can use equations of state \cite{KeT} \begin{equation} \omega_i = \frac{P_i}{\rho_i}, \end{equation} to integrate the energy conservation equation, for convenience written in the form \begin{equation} \label{eq:dotomega} \frac{\dot{\Omega_i}}{\Omega} = -3(1+\omega_i)\frac{\dot Q}{Q}, \end{equation} and relate the density parameter of each entity with the scale factor. This leads to the following behaviour. For radiation \begin{equation} \label{eq:3rad} \omega_r = \frac{1}{3},\hspace{1 cm} \Omega_r \propto Q(t)^{-4},\hspace{1 cm} T_r \propto Q(t)^{-1}; \end{equation} for matter (dust) \begin{equation} \label{eq:mat3} \omega_m = 0,\hspace{1 cm} \Omega_m \propto Q(t)^{-3},\hspace{1 cm} T_m \propto Q(t)^{-2}; \end{equation} and for dark energy \begin{equation} \omega_\Lambda = -1, \hspace{1 cm} \Omega_\Lambda \propto \text{constant}. \end{equation} | Through modelling the universe as a perfect fluid in terms of $Q(t)$, $\Omega(t)$ and $T(t)$, interpreting dark energy as arising from fluctuations of the metric and assuming neutralinos as dark matter particles, we could reconstruct many of the important events of its evolution. Starting from the known present epoch parameters, we show the temperature and density effects of particle annihilation and estimate the average present temperature of dark matter if it was composed of neutralinos.\\ We consider the two possibilities of Majorana and Dirac neutrinos. The overall development of the universe is not noticeably affected by the extra degrees of freedom of the Dirac neutrinos as compared to the Majorana case. We find the interesting result that at the epoch of primordial nucleosynthesis the sterile neutrino density, compared to the active neutrino density, is suppressed by a factor which is in agreement with estimates in the literature, when the sterile neutrinos start out, at the end of inflation, with the same temperature as the rest of the universe. | 14 | 4 | 1404.1833 |
1404 | 1404.1002_arXiv.txt | We present photospheric-phase observations of LSQ12gdj, a slowly-declining, UV-bright Type~Ia supernova. Classified well before maximum light, LSQ12gdj has extinction-corrected absolute magnitude $M_B = -19.8$, and pre-maximum spectroscopic evolution similar to SN~1991T and the super-Chandrasekhar-mass SN~2007if. We use ultraviolet photometry from \emph{Swift}, ground-based optical photometry, and corrections from a near-infrared photometric template to construct the bolometric (1600--23800~\AA) light curve out to 45~days past $B$-band maximum light. We estimate that LSQ12gdj produced $0.96 \pm 0.07$~\Msol\ of \nickel, with an ejected mass near or slightly above the Chandrasekhar mass. As much as 27\% of the flux at the earliest observed phases, and 17\% at maximum light, is emitted bluewards of 3300~\AA. The absence of excess luminosity at late times, the cutoff of the spectral energy distribution bluewards of 3000~\AA, and the absence of narrow line emission and strong \ion{Na}{1}~D absorption all argue against a significant contribution from ongoing shock interaction. However, $\sim 10$\% of LSQ12gdj's luminosity near maximum light could be produced by the release of trapped radiation, including kinetic energy thermalized during a brief interaction with a compact, hydrogen-poor envelope (radius $< 10^{13}$~cm) shortly after explosion; such an envelope arises generically in double-degenerate merger scenarios. | Type Ia supernovae (SNe~Ia) have become indispensable as luminosity distance indicators at large distances appropriate for studying the cosmological dark energy \citep{riess98,scp99}. They are believed to be the thermonuclear explosions of carbon-oxygen white dwarfs, and their spectra are generally very similar near maximum light, although some spectroscopic diversity exists \citep{bfn93,benetti05,branch06,branch07,branch08,wang09}. SNe~Ia used for cosmology are referred to as spectroscopically ``(Branch) normal'' \citep{bfn93} SNe~Ia; they have a typical absolute magnitude near maximum light in the range $-18.5 < M_V < -19.5$. They are used as robust standard candles based on empirical relations between the SN's luminosity and its colour and light curve width \citep{riess96,tripp98,phillips99,goldhaber01}. Maximum-light spectroscopic properties can also help to improve the precision of distances measured using normal SNe~Ia \citep{sjb09,wang09,csp10,fk11}. Another subclass of SNe~Ia with absolute magnitude $M_V \sim -20$ has also attracted recent attention. At least three events are currently known: SN~2003fg \citep{howell06}, SN~2007if \citep{scalzo10,yuan10}, and SN~2009dc \citep{yamanaka09,tanaka10,silverman11,taub11}. A fourth event, SN~2006gz \citep{hicken07}, is usually classed with these three, although its maximum-light luminosity depends on an uncertain extinction correction from dust in its host galaxy. These four events are spectroscopically very different from each other. SN~2006gz has a photospheric velocity typical of normal SNe~Ia as inferred from the velocity of the \ion{Si}{2} $\lambda 6355$ absorption minimum, and shows \ion{C}{2} absorption ($\lambda\lambda 4745, 6580, 7234$) in spectra taken more than 10~days before $B$-band maximum light. In contrast, SN~2009dc shows low \ion{Si}{2} velocity \vSi\ ($\sim 8000$~\kms), a relatively high \ion{Si}{2} velocity gradient \vdot\ ($\sim -75$~\kms~day$^{-1}$), and very strong, persistent \ion{C}{2} $\lambda6580$ absorption. SN~2007if is spectroscopically similar to SN~1991T \citep{filippenko92,phillips92} before maximum light, its spectrum dominated by \ion{Fe}{3} and showing only very weak \ion{Si}{2}, with a definite \ion{C}{2} detection in a spectrum taken 5~days after $B$-band maximum light. SN~2006gz, SN~2007if and SN~2009dc show low-ionization nebular spectra dominated by \ion{Fe}{2}, in contrast to normal SNe~Ia which have stronger \ion{Fe}{3} emission \citep{maeda09,taub13}. Only one spectrum, taken at 2~days past $B$-band maximum, exists for SN~2003fg, which resembles SN~2009dc at a similar phase. Recently two additional SNe, SN~2011aa and SN~2012dn, have been proposed as super-Chandrasekhar-mass SN~Ia candidates based on their luminosity at ultraviolet (UV) wavelengths as observed with the \emph{Swift} telescope \citep{brown14}. These extremely luminous SNe~Ia cannot presently be explained by models of exploding Chandrasekhar-mass white dwarfs, since the latter produce at most 1~\Msol\ of \nickel\ even in a pure detonation \citep{kmh93}. While they might more descriptively be called ``superluminous SNe~Ia'', these SNe~Ia have typically been referred to as ``candidate super-Chandrasekhar SNe~Ia'' or ``super-Chandras'', based on an early interpretation of SN~2003fg as arising from the explosion of a differentially rotating white dwarf with mass $\sim 2$~\Msol\ \citep{howell06}. Observation of events in this class has stimulated much recent theoretical investigation into super-Chandrasekhar-mass SN~Ia channels \citep{hachisu11,justham11,rds12,dm13a,dm13b}, and into mechanisms for increasing the peak luminosity of Chandrasekhar-mass events \citep{hsr07}. The status of superluminous SNe~Ia as being super-Chandrasekhar-mass has historically been closely tied to their peak luminosity. SN~2003fg's ejected mass was inferred at first from its peak absolute magnitude $M_V = -19.94$, requiring a large mass of \nickel\ \citep[$M_\mathrm{Ni} = 1.3 \pm 0.1$~\Msol;][]{arnett82} and a low \ion{Si}{2} velocity near maximum ($\sim 8000$~\kms), suggesting a high binding energy for the progenitor. Ejected mass estimates were later made for SN~2007if \citep{scalzo10} and SN~2009dc \citep{silverman11,taub11}, producing numbers of similar magnitude. These ejected mass estimates depend, to varying extents, on the interpretation of the maximum-light luminosity in terms of a large \nickel\ mass, which can be influenced by asymmetries and/or non-radioactive sources of luminosity. For example, shock interaction with a dense shroud of circumstellar material (CSM) has been proposed as a source of luminosity near maximum light for SN~2009dc \citep{taub11,hachinger12,taub13}. The CSM envelope would have to be largely free of hydrogen and helium to avoid producing emission lines of these elements in the shocked material. The additional luminosity could simply represent trapped radiation from a short interaction soon after explosion with a compact envelope, rather than an ongoing interaction with an extended wind. Such an envelope is naturally produced in an explosion resulting from a ``slow'' merger of two carbon-oxygen white dwarfs \citep{it84,shen12}. \citet{kmh93} modeled detonations of carbon-oxygen white dwarfs inside compact envelopes, calling them \emph{tamped detonations}; these events are luminous and have long rise times, but appear much like normal SNe~Ia after maximum light. A strong ongoing interaction with an extended wind, in contrast, is expected to produce very broad, ultraviolet (UV)-bright light curves and blue, featureless spectra uncharacteristic of normal SNe~Ia \citep{fryer10,bs10}. Searching for more candidate super-Chandrasekhar-mass SNe~Ia, \citet{scalzo12} reconstructed masses for a sample of SNe~Ia with spectroscopic behavior matching a classical 1991T-like template and showing very slow evolution of the \ion{Si}{2} velocity, similar to SN~2007if; these events were interpreted as tamped detonations, and the mass reconstruction featured a very rough accounting for trapped radiation. One additional plausible super-Chandrasekhar-mass candidate event was found, SNF~20080723-012, with estimated ejected mass $\sim 1.7 \Msol$ and \nickel\ mass $\sim 0.8 \Msol$. The other events either had insufficient data to establish super-Chandrasekhar-mass status with high confidence, or had reconstructed masses consistent with the Chandrasekhar mass. However, none of the \citet{scalzo12} SNe had coverage at wavelengths bluer than 3300~\AA, making it impossible to search for early signatures of shock interaction, and potentially underestimating the maximum bolometric luminosity and the \nickel\ mass. While \citet{brown14} obtained good UV coverage of two new candidate super-Chandrasekhar-mass SNe~Ia, 2011aa and 2012dn, no optical photometry redward of 6000~\AA\ has yet been published for these SNe, precluding the construction of their bolometric light curves or detailed inference of their masses. In this paper we present observations of a new overluminous ($M_B = -19.8$) 1991T-like SN~Ia, LSQ12gdj, including detailed UV (from \emph{Swift}) and optical photometric coverage, as well as spectroscopic time series, starting at 10~days before $B$-band maximum light. We examine the UV behavior as a tracer of shock interaction and as a contribution to the total bolometric flux, \revised{and perform some simple semi-analytic modeling to address the question: what physical mechanisms can drive the high peak luminosity in super-Chandrasekhar-mass SN~Ia candidates, and how might this relate to the explosion mechanism(s) and the true progenitor mass?} | 14 | 4 | 1404.1002 |
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1404 | 1404.0004_arXiv.txt | We present a method of frequency stabilizing a broadband etalon that can serve as a high-precision wavelength calibrator for an Echelle spectrograph. Using a laser to probe the Doppler-free saturated absorption of the rubidium $D_2$ line, we stabilize one etalon transmission peak directly to the rubidium frequency. The rubidium transition is an established frequency standard and has been used to lock lasers to fractional stabilities of $<10^{-12}$~\citep{Affolderbach,Ye1996}, a level of accuracy far exceeding the demands of radial velocity (RV) searches for exoplanets. We describe a simple setup designed specifically for use at an observatory and demonstrate that we can stabilize the etalon peak to a relative precision of $<10^{-10}$; this is equivalent to 3~cm/s RV precision. | Detecting an earthlike planet in the habitable zone of a solar-type star requires a radial velocity (RV) precision of a few cm/s, an order of magnitude better than the best currently attainable precision, which is $\sim70$~cm/s, of which 30~cm/s is due to the wavelength calibration uncertainty~\citep{Lovis2006, Dumusque2013}. To detect a velocity change of 3~cm/s, a relative wavelength shift of $v/c=10^{-10}$ must be measured at a typical spectrograph resolution of only $10^5$. Detecting a shift of only $10^{-5}$ of a resolution element is an extraordinary challenge, and the calibration has to track changes in the spectrograph on that level. Currently available calibration technologies using thorium-argon (ThAr) emission lamps or gas absorption cells are not sufficiently precise for this. With a range of new spectrographs under construction or coming into operation over the next 5-10 years~(ESPRESSO~\citep{Espresso}; CODEX~\citep{Codex}; NRES~\citep{Nres}; IRD~\citep{Ird}; CARMENES~\citep{Quirrenbach2010}; HPF~\citep{Hpf}; HRS~\citep{Hrs}; SPIRou~\citep{Spirou}), there is clearly demand for a reliable and affordable calibration method with intrinsic precision and long-term stability in the cm/s range. Two techniques currently being explored for producing calibration spectra that meet this requirement are laser frequency combs (LFCs) and Fabry-Perot etalons. While LFCs promise to exceed the requirements on precision and stability, they are very complex systems with a high cost of ownership. A broadband Fabry-Perot etalon illuminated with white light produces a spectrum of regularly spaced lines, creating a ``passive comb" that is far simpler and less expensive than a LFC; however, the etalon does not provide an absolute frequency reference and must be stabilized to be useful for calibration. Nevertheless, impressive results have been achieved using passively stabilized etalons for spectrograph calibration~\citep{Wildi2012}. In this paper, we describe a method of stabilizing a Fabry-Perot etalon and referencing it to a known absolute frequency by locking it to an atomic transition. We use saturated absorption spectroscopy to stabilize one transmission peak of a broadband etalon to a hyperfine transition of rubidium. The rubidium transition is widely used as a frequency standard, which has been utilized to stabilize lasers to relative precisions of $<10^{-12}$~\citep{Ye1996,Affolderbach}; it provides an ideal, repeatable, and stable wavelength reference. Our locking scheme is simple and robust and will work for a wide variety of etalons operating in the visible and near-infrared (NIR). | \subsection{Reliability} Reliability, the necessary amount of maintenance, and cost of operation are important considerations for subsystems of a spectrograph installed at an observatory. With this in mind, we chose commercial off-the-shelf components for parts that could break. The laser diode is probably the most likely component to fail. However, a mean time between failures of up to 100,000 hours, or $>10$ years of continuous operation, is reported~\citep{ILX33}. We run the laser diode at 50\% of the rated maximum current and cool it below ambient temperature, both of which improve lifetime. Similarly, the PZT that scans the laser is used far below its maximum voltage rating. Instead of biasing the PZT, we use a fine-pitched screw to adjust the frequency of the laser to the Rb line. Since we scan fairly slowly, the power dissipation inside the PZT is low. Lastly, we smooth out the sawtooth applied to the PZT to avoid stress at the turning points. Changing the laser diode or PZT is straightforward. Maintenance of the laser head does not introduce any offset when the lock is resumed, since the frequency zero point is derived directly from the Rb transition. Several external parameters can affect the observed Rb transition frequency in a saturation spectroscopy setup like ours and must be carefully controlled. The most important of these are pump beam intensity, angle between the pump and probe beams, Rb cell temperature, and ambient magnetic field. \citet{Affolderbach} measured the line shifts due to these parameters using a setup similar to ours and demonstrated sufficient control of all of them to achieve $< 2 \times 10^{-12}$ relative frequency stability of an ECDL locked to the Rb $D_2$ line over timescales $\ge10^4$~s. Controlling all above-mentioned shifts sufficiently well to guarantee stability and reproducibility of the Rb frequency at the level of precision needed for calibration is relatively straightforward, for example by surrounding the Rb cell with a single layer of mu-metal and actively stabilizing the cell temperature and pump laser power. The most critical part of the setup is the etalon itself. A replacement etalon will have a slightly different line spacing and probably a different dispersion. It will need to be calibrated against an absolute frequency standard that covers the entire bandwidth, ideally a laser frequency comb, and this calibration compared to the one of the original etalon. However, breaking the etalon assembly seems unlikely. For long term stability, a thermally adjusted etalon seems to be a more robust choice, as in this case the cavity is completely passive, while breaking the PZT in a PZT-driven etalon would render the etalon unusable. Finally, metallic mirror coatings on the etalons could change over time. The protected silver coatings we employ can degrade through corrosion. For bulk etalons operated inside a vacuum housing, this is unlikely. Noncorroding gold coatings can be used for etalons in the NIR. The coating of an FFP experiences mechanical stress since the cavity fiber is in physical contact with the input and output fibers. Also, depending on the power of the white light source and laser, the flux at the mirrored ends of the single-mode waveguide that forms the etalon can be very high, potentially damaging the coating. Very dense, low loss dielectric mirror coatings are likely better equipped to withstand these conditions. \subsection{Future work} Current work is focused on integrating the system into a compact and rugged unit that can be used at an observatory. This includes improved shielding against electrical interference, improved thermal isolation of the setup, and enclosing the Rb cell in a magnetic shield. To reduce the laser's sensitivity to mechanical vibrations and thermal fluctuations, we plan to implement a simple yet effective active stabilization scheme based on polarization spectroscopy~\citep{Fuehrer2}. We will also optimize the algorithm we use to determine the Rb reference frequency, which we believe is a major contribution to our error budget. We plan to record simultaneous ThAr and laser-locked FPE spectra with our $R=80,000$ laboratory Echelle spectrograph to map out the dispersion of the etalon and determine the wavelength of each peak. The precision of this measurement will be limited by the precision of the ThAr calibration. The stability of the entire etalon spectrum can be verified with a precision equal to that of our locking technique by simultaneously measuring a second etalon line using a second laser and another atomic transition, for instance the cesium line at 852~nm. Another possibility is frequency doubling a 1560~nm laser and using the second harmonic at 780~nm in the same way we use our laser~\citep{Masuda2007}, but also using the original wavelength to monitor an etalon line at 1560~nm, which would work well for NIR etalons. For etalons in the visible, the same thing can be accomplished by frequency doubling a 780~nm laser and monitoring an etalon line at 390~nm in addition to the original 780~nm line used for locking. \subsection{Summary} In summary, we have presented a simple setup that reliably locks an etalon suitable for calibrating an Echelle spectrograph with better than 3~cm/s radial velocity precision for any realistic exposure time. Care was taken to adapt the locking scheme to the specific requirements of an observatory. We use a minimum of optical and electronics hardware and use a single locking loop that references the etalon directly to the rubidium transition. This technique can be used to stabilize a variety of etalons in the visible and NIR, and we have demonstrated excellent precision with thermal as well as PZT-driven locking of air-spaced and fiber-based etalons. Our setup combines long lifetime and high reliability with low cost. It can be easily integrated with any of the etalons currently investigated for astronomical use. Even if a particular etalon cannot be locked to the rubidium transition (i.e. due to a very long time constant), our scheme allows precise real-time monitoring of the etalon drift relative to the rubidium. In this way, observations calibrated using the etalon can be accurately corrected for etalon drift. A laser-locked etalon provides high precision, long-term stability, and a wide spectral range for calibration, making it an excellent, cost-effective solution for calibrating an Echelle spectrograph. | 14 | 4 | 1404.0004 |
1404 | 1404.1880_arXiv.txt | We propose an economical model in which a singlet Z$_2$-odd scalar field accounts for the primordial inflation and the present dark matter abundance simultaneously in the light of recent BICEP2 result. Interestingly, the reheating temperature and the thermal dark matter abundance are closely connected by the same interaction between the singlet scalar and the standard model Higgs. In addition, the reheating temperature turns out to be quite high, $T_\text{R} \gtrsim 10^{12}\,\GEV$, and hence the thermal leptogenesis is compatible with this model. Therefore, it can be one of the simplest cosmological scenarios. | \label{sec:} Recently, the BICEP2 experiment discovered the B-mode polarization in the cosmic microwave background (CMB) anisotropy, which is interpreted as the primordial gravitational waves of the inflationary origin~\cite{Ade:2014xna}. This confirms the idea of inflation~\cite{Guth:1980zm,Linde:1981mu}, especially the high scale inflation such as the chaotic inflation~\cite{Linde:1983gd}. On the other hand, one of the greatest mysteries of the Universe is the presence of dark matter (DM)~\cite{Ade:2013rta}. Since there is no candidate for the DM in the standard model (SM) of particle physics, it clearly requires physics beyond the SM. Maybe the simplest extension of the SM is to add a singlet scalar field $\phi$ which has a Z$_2$-symmetry~\cite{Silveira:1985rk,McDonald:1993ex} and couples to the SM Higgs boson $H$ in the scalar potential as \begin{equation} V = \frac{1}{2}m_\phi^2\phi^2 + \frac{1}{2}g^2\phi^2|H|^2, \end{equation} where $g$ is a coupling constant. Due to the Z$_2$-symmetry under which $\phi$ transforms as $\phi \to -\phi$, it is stable. It can have a correct annihilation cross section through the Higgs portal for $m_\phi \sim \mathcal O(100)$\,GeV and $g \sim \mathcal O(1)$ and account for the observed amount of DM. We show that the scalar singlet DM, $\phi$, can cause inflation which is consistent with the BICEP2 result. Naively, at the large field value, $\phi$ obtains a $\phi^4$ potential radiatively, and hence this chaotic inflation with $\phi^4$ potential is already ruled out. Moreover, it is difficult to account for the observed density perturbation of the Universe. Our idea is to modify the kinetic term of $\phi$ so that the potential becomes quadratic in terms of the canonically normalized field. This is the so-called running kinetic inflation~\cite{Nakayama:2010kt,Nakayama:2010sk}. It has been shown that the SM Higgs boson can be the inflaton to be consistent with the BICEP2 result~\cite{Nakayama:2014koa}. In this paper, we identify the singlet scalar $\phi$ as the inflaton, and show that it can simultaneously explain the present DM abundance. The key feature is that the inflaton coherent oscillation can be soon dissipated away even if it is perturbatively stable at its vacuum, and eventually the inflaton itself participates in the thermal plasma as extensively studied in Refs.~\cite{Mukaida:2012qn,Mukaida:2013xxa,Mukaida:2014yia}. Then, as the Universe expands, the annihilation of the inflaton particles again decouples from the thermal plasma at late time, which leads to the standard freeze-out DM production. This scenario typically results in a high reheating temperature $T_\text{R} \gtrsim 10^{12}\, \GEV$ that is compatible with the thermal leptogenesis~\cite{Fukugita:1986hr}. Interestingly, the coupling $g$, which determines the present relic DM abundance, also determines the reheating temperature of the Universe. In this sense, the model we propose is quite economical:\footnote{ A similar model was proposed in Refs.~\cite{Lerner:2009xg,Okada:2010jd} in the context of inflation with non-minimal coupling to gravity~\cite{Bezrukov:2007ep}. In Ref.~\cite{Nakayama:2010kt}, the possibility of inflatino DM (the supersymmetric partner of the inflaton) in the context of running kinetic inflation was pointed out. } we can explain the inflation, reheating and DM consistent with observations by just adding a real scalar $\phi$. | \label{sec:} Motivated by the recent observation of the B-mode polarization by the BICEP2 experiment, we have considered a scenario that the chaotic inflation is induced by a singlet scalar field with the running kinetic term, and it simultaneously becomes DM in the present Universe. The key point is that the inflaton is heavy at the large field value where inflation happens, while it can be so light around the potential minimum that its thermal relic abundance can match with observed DM abundance. The process of reheating might be non-trivial since the inflaton has a Z$_2$-symmetry and it is perturbatively stable at its vacuum. However, the combined effects of particle production and the scattering with particles in thermal bath cause efficient dissipation on the inflaton coherent oscillation and actually the inflaton is soon thermalized after inflation ends. Once the inflaton participates in the thermal plasma, the following thermal history is described by the standard radiation dominated Universe. Since the reheating temperature is so high, $T_\text{R} \gtrsim 10^{12}\, \GEV$, the thermal leptogenesis successfully works~\cite{Fukugita:1986hr}. Interestingly, the four point interaction with the SM Higgs that determines the DM relic density also induces the cosmic reheating and determines the reheating temperature. Thus we think that this is a kind of minimal scenario that explains the primordial inflation, reheating and present DM abundance. The singlet scalar DM scenario may be probed by future direct DM detection experiments~\cite{Aprile:2012zx} and also by the collider searches through the Higgs invisible decay~\cite{Peskin:2012we}. Because of the high inflation scale and high reheating temperature, direct detection of inflationary gravitational waves with future space laser interferometers is also plausible~\cite{Turner:1990rc,Seto:2003kc,Smith:2005mm,Nakayama:2008wy}. | 14 | 4 | 1404.1880 |
1404 | 1404.4188_arXiv.txt | The positron flux measured near Earth shows a rise with energy beyond 30 GeV. We show that this rise may be compatible with the production of positrons in $p \gamma$ interactions in the jets of microquasars. | The electron and positron spectra measured by cosmic ray experiments (Boezio et al. 2000, DuVernois et al. 2001, Aguilar et al. 2013, Adriani et al. 2013, Ackermann et al. 2012) are interesting probes of Nature. Electrons and positrons are injected in our Galaxy by cosmic-ray sources and interactions during the propagation of cosmic ray protons and nuclei. The electron flux times the electron's energy cube (e.g., in units of GeV$^2$ m$^{-2}$ s$^{-1}$ sr$^{-1}$) is a useful representation {\bf of the spectrum}. It is nearly constant in the energy range $10 $ GeV -- $1 $ TeV, and steeply falls down at higher energies. The flux measured up to 4 TeV is composed of primary electrons and secondaries from various interactions. Production of pairs in cosmic ray interactions with matter and radiation, electrons emitting pairs in a magnetic field $(e \, B\rightarrow e^+ \ e^-)$, and $\gamma \, \gamma\rightarrow e^+ \ e^-$ interactions contribute to the electron and positron fluxes equally. However, the recent measurements of the positron flux shows a rise beyond 30 GeV, which could have its origin either in conventional astrophysical sources or in dark matter annihilations, see e.g., Finkbeiner (2011), Gaggero et al. (2013), Cholis \& Hooper (2013), Mauro et al. (2014), Mertsch \& Sarkar (2014). A fraction of the positron flux measured up to 350 GeV originates from the pair-producing interactions mentioned above. Photo-hadronic interactions leading to the production of positrons $(p \,\gamma\rightarrow \Delta^+\rightarrow \pi^+ \, n$, $\pi^+ \rightarrow e^+ \, \nu_{\mu} \, \bar\nu_{\mu} \, \nu_e )$ may also contribute to the observed positron spectrum; see, e.g., Gupta \& Zhang (2008). In the present work, we discuss the spectrum of positrons expected from these interactions in the jets of microquasars and how this excess flux may help explaining the recent measurements by PAMELA (Aguilar et al. 2013), AMS02 (Adriani et al. 2013) and Fermi experiments (Ackermann et al. 2012). Microquasars were considered earlier to explain the positron annihilation radiation at 511 keV (Guessoum, Jean \& Prantzos 2006, Vila \& Romero 2010). The possibility of explaining the observed positron excess at tens of GeV with microquasars has been mentioned in the review by Fan et al. 2010, but up to now, no computation that demonstrates that this is indeed feasible nor any precision on which process would make this excess happen is available in literature. Here we use the positron data to study this possibility quantitatively. The aim of this paper is not to provide a model of the inner MQ engine (and thus to construct a detailed SED of the photons emitted by it), which has been done in several papers, with different levels of detail. There are several situations where the $p\gamma$ process dominates over leptonic and other hadronic interactions (pp) at high energies. For instance see model A, C or D in the paper by Vila et al. 2012. In these models, the target radiation field considered for the $p\gamma$ interactions are the synchrotron photons of primary electrons and a detailed description of the location of the acceleration region and the magnetic field dependence along the jet axis can be found. | The recent detection of Doppler-shifted X-ray emission lines from a typical black-hole candidate X-ray binary, 4U~1630--47, coincident with the reappearance of radio emission from the jets of the source, implies that baryons can be accelerated in jets of microquasars \citep{diaz}. The jets should be strong sources of gamma-rays and neutrinos, and in principle could contribute to the observed positron excess significantly. The positron excess has been studied earlier with $pp$ interaction models, see e.g., Gaggero et al. (2013), and constrained with observed $B/C$ ratios, see, e.g., Cholis \& Hooper (2013), Mertsch \& Sarkar (2014). Here, we have shown that $p\gamma$ interactions in boosted environments such as jets of microquasars may help in explaining the observed rise in the positron spectrum beyond 30 GeV. Low mass microquasars are of special interest in this regard, for hadronic models based on inelastic $pp$ collisions are not expected to play a leading role, the companion star being cold and old. Because of the same reasons, the external photon background are scarce, what would limit (albeit not rule out in all cases, especially due to self-synchrotron Compton, see, e.g., Bosch-Ramon et al. 2006a,2006b) the ability of leptonic processes to dominate the spectra. If jets accelerate protons, these sources may lead to multi-particle injection via $p\gamma$ processes, where perhaps the photons are synchrotron generated at the base of the jets, see e.g., Levinson \& Waxman 2001. For recent models of proton low-mass microquasars see e.g., Romero \& Vila 2008, 2010; Saitou et al. 2011; Zhang et al. 2010; Vila et al. 2012. Luminosities of the $p\gamma$ channel in these works for individual sources are in agreement with the requirements found in our work in order to explain a significant part of the positron excess with microquasar jets. We finally compare our scenario of $p\gamma$ interactions and subsequent photo-pion decay with the scenario of cosmic ray interactions in the interstellar medium discussed by Cowsik, Burch \& Maziwa-Nussinov (2013). In this paper the authors have interpreted the observed positron spectrum as the secondaries produced in interactions of primary cosmic ray nuclei with interstellar medium assuming the positrons stay for 2 Myr in the Galaxy. According to their prediction the positron spectrum is proportional to $ E_{\rm{e}}^{-3.65}$ above 300 GeV. In our case the spectrum is proportional to $E_{\rm{e}}^{-2.8}$ up to the break energy in the positron spectrum at 630 GeV if the break energy ($\epsilon_{\rm{br}}$) in the synchrotron spectrum of photons from the jets is at 0.1 MeV. Above the break energy the spectral index of the positron spectrum would be proportional to $E_{\rm{e}}^{\gamma_1-4-1/2}$ where $\gamma_1$ is the spectral index of the photon spectrum below $\epsilon_{br}$. In some cases, the positron flux from these sources could be higher than their electron flux, for instance, if their luminosity in primary protons is higher than in electrons. If so, we expect these sources to be positron dominated and not contribute significantly to the observed diffuse electron flux. | 14 | 4 | 1404.4188 |
1404 | 1404.1078_arXiv.txt | { Many classes of active galactic nuclei (AGN) have been defined entirely throughout optical wavelengths while the X-ray spectra have been very useful to investigate their inner regions. However, optical and X-ray results show many discrepancies that have not been fully understood yet. } {The main purpose of the present paper is to study the ``synapses'' (i.e., connections) between the X-ray and optical AGN classifications. } {For the first time, the newly implemented {\sc efluxer} task allowed us to analyse broad band X-ray spectra of a sample of emission line nuclei without any prior spectral fitting. Our sample comprises 162 spectra observed with \emph{XMM}-Newton/pn of 90 local emission line nuclei in the Palomar sample. It includes, from the optical point of view, starbursts (SB), transition objects (T2), low ionisation nuclear emission line regions (L1.8 and L2), and Seyfert nuclei (S1, S1.8, and S2). We use artificial neural networks (ANNs) to study the connection between X-ray spectra and the optical classes. } {Among the training classes, the ANNs are 90\% efficient at classifying the S1, S1.8, and SB classes. The S1 and S1.8 classes show a negligible SB-like component contribution with a wide range of those from S1- and S1.8-like components. We suggest that this broad range of values is related to a large degree of obscuration in the X-ray regime. When including all the objects in our sample, the S1, S1.8, S2, L1.8, L2/T2/SB-AGN (SB with indications of AGN activity in the literature), and SB classes have similar average X.ray spectra, but these average spectra can be distinguished from class to class. The S2 (L1.8) class is linked to the S1.8 (S1) class with larger SB-like component than the S1.8 (S1) class. The L2, T2, and SB-AGN classes conform a class in the X-rays similar to the S2 class albeit with larger fractions of SB-like component. We argue that this SB-like component might come from the contribution of the host galaxy emission to the X-rays, which is large when the AGN is weak. Up to 80\% of the emission line nuclei and, on average, all the optical classes included in our sample show a non-negligible fraction of S1-like or S1.8-like component. Thus, an AGN-like component seems to be present in the vast majority of the emission line nuclei in our sample. } {The ANN trained in this paper is not only useful to study the synergies between the optical and X-ray classifications, but also could be used to infer optical properties from X-ray spectra in surveys like \emph{eRosita}.} | \label{sec:introduction} At optical wavelengths emission line galaxies can be grouped into HII nuclei, active galactic nuclei (AGN), galaxies with low-ionisation nuclear emission line regions (LINERs), and transition objects \citep[whose optical spectra are intermediate between those of pure LINERs and HII regions; see][for a review]{Ho08}. Optical spectroscopic studies have shown that only 10\% of nearby galaxies are Seyferts, while LINERs and transition objects account to no more than 20\% and 10\% of them, respectively \citep[e.g., Palomar Survey by][]{Ho97}. HII nuclei are powered by a compact star forming region. In AGN, the major energy source is assumed to be accretion of matter into a super-massive black hole (SMBH). The nature of the main energy source in LINERs (and transition objects) is not settled yet. They might be low-luminosity AGN (LLAGN), in which case, they will constitute the main fraction of the AGN population \citep{Heckman80,Ho97}. However, other emission mechanisms like shock heating \citep{Dopita95}, OB stars in compact nuclear star clusters \citep{Terlevich85}, or pre-main sequence stars ionisation \citep{Cid-Fernandes04} have also been proposed. AGN are traditionally divided into two main classes, namely Type-1 and Type-2 objects, based on the existence (Type-1) or not (Type-2) of broad permitted lines (FWHM$\rm{>}$2000 km $\rm{s^{-1}}$). The so-called unification model (UM) proposes that both types of AGN are essentially the same objects viewed at different angles \citep{Antonucci93,Urry95}. An optically thick dusty torus surrounding the central source would be responsible for blocking the region where these broad emission lines are produced (the broad line region, BLR) in Type-2 Seyferts. Therefore, Type-2 Seyferts are essentially Type-1 Seyferts blocked by the dusty torus along the line of sight (LOS) to the observer. A strong observational evidence in favour of a unification between Type-1 and Type-2 Seyferts was the discovery of broad optical lines in the polarised spectrum of the archetypal Type-2 Seyfert, NGC\,1068 \citep{Antonucci85}. The torus must not be spherically symmetric, in order to obscure the BLR, allowing at the same time the region producing the permitted narrow lines (known as narrow-line region, NLR) to reach us from the same LOS. The locus of this obscuring material was initially postulated at parsec scales and confirmed by modelling the spectral energy distribution (SED) of Seyferts \citep[e.g.,][]{Ramos-Almeida11,Alonso-Herrero11} and by interferometric observations \citep[e.g., Circinus galaxy,][]{Tristram07}. Such scales are unreachable with the current instrumentation, so the torus morphology can only be inferred by indirect measurements. Although the UM is widely accepted for many classes of Seyferts, there is still no consensus on its general applicability for all members of each class \citep[see][for a review]{Bianchi12}. An example of this mismatch is the so-called `optically elusive' AGN \citep{Maiolino98}. These elusive AGN are nuclear hard X-ray sources whose intrinsic luminosities are in the Seyfert range but they lack of optical Seyfert-like signatures. Another example is that about half of the brightest Type-2 Seyferts are characterised by the lack of BLR even with high-quality spectro-polarimetric data \citep[known as `True Type-2' Seyferts,][]{Tran01,Tran03}. These Type-2 Seyferts without BLR are expected to occur theoretically at low accretion rates or low luminosities \citep{Elitzur09}. With respect to LINERs, even if they are powered predominately by accretion into a SMBH, it is unclear whether the UM can also apply to these LLAGN. Indeed, both a different accretion mode and large amounts of obscuration have been proposed to explain the differences between LINERs and Seyferts \citep{Gonzalez-Martin06,Gonzalez-Martin09a,Gonzalez-Martin09b,Younes11,Hernandez-Garcia13}. X-rays in AGN are thought to originate in the innermost region of the accretion flow and are also thought to be affected by the obscuring material along the LOS. X-ray observations of AGN have provided additional evidence in favour of the UM. For example, the obscuring material along the LOS (measured at X-rays by the hydrogen column density, $\rm{N_H}$) is substantially larger in Type-2 Seyferts than in Type-1 Seyferts \citep[e.g.,][]{Maiolino98,Risaliti99,Panessa06,Cappi06}. Although modelling of X-ray spectra is one of the best ways to estimate the obscuration, it has also some caveats. For example, the obscuration measured in Seyferts depends on the model used for the underlying X-ray continuum. The main aim of this paper is to investigate if objects in different (optical) classes have similar X-ray spectra, and if they do, whether their average X-ray spectrum differs between the different classes or not. Furthermore, we compare the average X-ray spectra of these classes in a model independent way. Consequently, instead of fitting each individual spectrum with a suitable model, we chose to use artificial neural networks (ANNs). We have selected for our analysis the X-ray spectra of 90 well-classified emission line nuclei included in the optically classified sample of nearby galaxies presented by \citet{Ho97}. We used ANNs to classify their X-ray spectrum and compare the average spectra of each class, without any model pre-assumptions. The main questions we address in this paper are the following: (1) how do optical classes ``behave'' at X-rays? in other words, do objects of the same (optical) class have the same X-ray spectrum (on average), and if yes, are the average X-ray spectra of the various optical classes the same or not? (2) If they are different, can we understand what is the main physical parameter that drives those differences? and (3) are AGN-like nuclei present in all emission line nuclei in nearby galaxies? does this include those galaxies that have absent or weak AGN signatures at optical wavelengths? Section \ref{sec:sample} gives the details on the selected sample and Section \ref{sec:reduction} the technical details of the reduction process. In Section \ref{sec:ANN} we describe the methodology and the main results of the ANN are presented in Section \ref{sec:ANNresults}. These results are discussed in Section \ref{sec:discussion} and summarised in Section \ref{sec:summary}. Along the paper a value of $\rm{H_{0} = 75}$ km s$^{-1}$ Mpc$^{-1}$ is assumed. | \label{sec:discussion} We have shown that the ANN analysis can be useful to classify the main optical classes using only X-ray spectra. In general, an object with $\rm{\nu_{SB}\le10}$ is almost certainly a S1 or a S1.8. Moreover, an object with little $\rm{\nu_{S1.8}}$ and large $\rm{\nu_{S1}}$ and $\rm{\nu_{SB}}$ is most probably a L1.8, while an object with little $\rm{\nu_{S1}}$ and large $\rm{\nu_{S1.8}}$ and $\rm{\nu_{SB}}$ is most probably a S2. Larger fractions of $\rm{\nu_{SB}}$ characterise the L2,T2 and SB nuclei. However, we would like to stress that most of the differences are found when we consider the average value for each class. Thus, although we believe that the ANN method is very useful to study the average properties, it may not be as successful in classifying a single object based on its ANN components. Using the results regarding the average properties of the objects in each class, in this section we discuss the following questions: (1) Type-1/Type-2 dichotomy; (2) optical versus X-ray classes; and (3) elusive AGN. Finally, we present the utility of this analysis for its application to X-ray surveys. \begin{figure} \centering \includegraphics[width=1.\columnwidth]{MeanANN3vsLsteep.png} \caption{The logarithmic of the 2-10 keV band observed luminosity, log(L(2-10 keV)), versus the $\rm{\nu_{S1}}$ (left) and $\rm{\nu_{S1.8}}$ (right) components. (Bottom): The logarithmic of the ratio between the observed luminosity at 6 keV versus the observed luminosity at 2 keV, $\rm{log(L_{6keV}/L_{2keV})}$, versus the $\rm{\nu_{S1}}$ (left) and $\rm{\nu_{S1.8}}$ (right) components. We only plot objects with $\rm{\nu_{SB}<10}$ (see text). The optical classes are shown as: S1 (red up-side down triangles), S1.8 (orange triangles), S2 (yellow diamonds), and L1.8 (purple stars).} \label{fig:MeanANN3vsLsteep} \end{figure} \subsection{Type-1/Type-2 dichotomy}\label{sec:dichotomy} Our results indicate that the X-ray spectra of the S1 and S1.8 classes can be reproduced by a mixture of the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ components, with $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ being stronger in the former and latter classes, respectively. Furthermore, the S1 and S1.8 classes show a continuous range of values of the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ components (see Fig. \ref{fig:ANNnuvsnu}, top-left panel). Our analysis cannot offer direct indications of the nature of the $\rm{\nu_{S1}}$ or the $\rm{\nu_{S1.8}}$ components, or for the physical parameter that drives their correlation for S1s and S1.8s. Below we discuss possible interpretations of this result. The continuous range of values for $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ could reflect a continuous range of absorptions (i.e., $\rm{N_H}$), increasing for the S1.8 class. This is consistent with the UM of AGN. Indeed, X-rays have been used in AGN to study the amount of absorption \citep{Risaliti99,Bianchi12,Ho08}. \citet{Risaliti99} found that 75\% of their Type-2 Seyferts were heavily obscured ($\rm{N_{H}>10^{23}cm^{-2}}$), 50\% of them were Compton-thick (i.e., $\rm{N_{H}>1.5\times 10^{24}cm^{-2}}$), with the S1.8 class characterised by an average lower $\rm{N_H}$ than the S2 class. Alternatively, a low flux level continuum is recently suggested by \citet{Elitzur14} as the main reason to classify objects as S1.8s. They suggest that intermediate types of objects are part of an evolutionary sequence where the BLR slowly disappears as the bolometric luminosity decreases. Hence, the continuous range of values for $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ could be interpreted either as (1) an increase of the absorption as we move from S1s and S1.8s or (2) a decrease of the AGN continuum flux in S1.8s. As shown below, our results favour the first interpretation. Assuming that L(2-10 keV) is an indication of the total luminosity, we would expect it to be proportional to $\rm{\nu_{S1}}$ and inversely correlated with $\rm{\nu_{S1.8}}$ if a decrease of the intrinsic continuum is responsible for the S1.8 class. Fig. \ref{fig:MeanANN3vsLsteep} (top panels) shows the log(L(2-10 keV))\footnote{L(2-10 keV) is computed as the sum of all the bins in the calibrated spectra in the 2-10 keV band multiplied by the size of the spectral bin ($\rm{\Delta E = 0.05}$ keV).} versus $\rm{\nu_{S1}}$ (left) and $\rm{\nu_{S1.8}}$ (right) for objects with a negligible contribution of $\rm{\nu_{SB}}$ ($\rm{\nu_{SB}<10}$). At each $\rm{\nu_{S1}}$ or $\rm{\nu_{S1.8}}$ values there is a large scatter of luminosities, but objects with large (small) $\rm{\nu_{S1}}$ ($\rm{\nu_{S1.8}}$) have larger X-ray luminosities, on average. The Pearson's correlation coefficients are r=0.37 ($\rm{P_{null}=0.008}$) and r=0.34 ($\rm{P_{null}=0.015}$) for the correlations with $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$, respectively (see Fig. \ref{fig:MeanANN3vsLsteep}, top, right and left panels). The small numbers of the correlation coefficients shows that the correlations are not strong, although the null hypothesis probability indicates that it may be significant. The bottom panels of Fig. \ref{fig:MeanANN3vsLsteep} show the steepness of the spectra, expressed as $\rm{log(L_{6keV}/L_{2keV})}$\footnote{$\rm{L_{2keV}}$ and $\rm{L_{6keV}}$ are the monochromatic luminosities at 2 keV and 6 keV, respectively, obtained from the flux-calibrated spectra.}, versus the $\rm{\nu_{S1}}$ (left) and $\rm{\nu_{S1.8}}$ (right) components. The X-ray spectra become harder (i.e., the emission at 6 keV becomes more prominent compared to the emission at 2 keV) when the $\rm{\nu_{S1.8}}$ component increases (and $\rm{\nu_{S1}}$ decreases). The correlation between them shows Pearson's correlation coefficients and null probabilities of r=0.91, $\rm{P_{null}=6\times10^{-20}}$ and r=0.89, $\rm{P_{null}=7\times10^{-18}}$, respectively. Irrespective of the reason for the spectral hardening, the strength of the correlations in the lower panels of Fig. \ref{fig:MeanANN3vsLsteep} indicates that the distributions of the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ components in Seyferts are not driven, primarily, by luminosity, but by the spectral hardening of their X-ray spectra. The simplest explanation for this spectral hardening is an increase of absorption, which in the case of Compton-thin sources affects much stronger the 2 keV flux than the 6 keV flux. Therefore, based on the strength of the correlations shown in Fig. \ref{fig:MeanANN3vsLsteep} it seems reasonable to assume that a variable amount of obscuration is the main physical parameter responsible for the continuous range of $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$. The same effect can also explain the weak correlations with the luminosity (see Fig. \ref{fig:MeanANN3vsLsteep}, top panels). If the observed luminosities are corrected for absorption, then both S1.8 and S1 could show the same level of X-ray luminosity. Therefore, we believe that the scenario to be preferred is that in which obscuration is responsible for the Type-1/Type-2 dichotomy. This is fully consistent with the UM of AGN, in which the obscuring torus is responsible for blocking the inner parts of the AGN (both the BLR and the X-ray source) in Type-2 galaxies. \begin{figure} \centering \includegraphics[width=1.\columnwidth]{NHvsANN.png} \caption{The logarithmic of the $\rm{N_H}$ versus $\rm{log(\nu_{S1.8}+20)}$. Filled symbols show $\rm{N_H}$ values using a simple power-law model to the 2-10 keV band (see Appendix A). Empty symbols show $\rm{N_H}$ reported in the literature when available (see Table \ref{tab:NHs}). Dashed vertical lines link the $\rm{N_H}$ values using a simple power-law model and those reported in the literature.} \label{fig:NHvsANN} \end{figure} A final check on the nature of this dichotomy can be performed comparing $\rm{\nu_{S1.8}}$ with the absorbing column density, $\rm{N_H}$, for these observations (see Fig. \ref{fig:NHvsANN} and Appendix A for the details on the measurements of $\rm{N_H}$). The quantity $\rm{log(\nu_{S1.8}+20)}$\footnote{Note that we have computed the logarithmic of $\rm{(\nu_{S1.8}+20)}$ in order to avoid negative values of $\rm{\nu_{S1.8}}$.} is linearly related with $\rm{log(N_H)}$ (r=0.93, $\rm{P_{null}=1.5\times10^{-21}}$) when derived with a simple power-law fit (filled symbols in Fig. \ref{fig:NHvsANN}). A less significant linear relation (r=0.57, $\rm{P_{null}=5.3\times10^{-3}}$) is found when using $\rm{N_H}$ estimates reported in the literature (empty symbols in Fig. \ref{fig:NHvsANN}). We believe this weaker relationship is due to: (1) less number of observations with $\rm{N_H}$ and (2) different models used for the spectral fittings for each observation. It reinforces the importance on the do a self-consistent modelling for the sample to compare the parameters. \subsection{Optical versus X-ray classes}\label{sec:BPT} The ANN has found differences on the average X-ray spectra of the six different classes: S1, S1.8, S2, L1.8, L2/T2/SB-AGN, and SB. Thus, the L2, T2, and SB-AGN belong to the same X-ray category according to the ANN results. It is worth to remark that division lines in the BPT diagrams have been developed and adapted as a function of the ionisation models and/or observations available \citep[e.g.,][]{Veilleux87,Osterbrock89,Kewley01,Kauffmann03,Kewley06,Stasinska06,Kewley13}. Objects close to the division between star-forming galaxies and AGN could be classified as L2, T2, or SB depending on how these divisions are set and/or how these three diagrams are used together. This could explain why the L2, T2, and SB-AGN classes cannot be distinguished at X-rays according to the ANN. Alternatively, the number of physical parameters governing the classes at X-rays might be lower than those driving the optical classes. \begin{figure} \centering \includegraphics[width=1.\columnwidth]{MeanANN3vsLhard.png} \caption{The mean $\rm{\overline{\nu}_{SB}}$ component versus the mean 2-10 keV band observed luminosity in logarithmic scale, $\rm{log(\overline{L}(2-10~keV))}$, per optical class. The optical classes are shown as: S1 (red up-side down triangles), S1.8 (orange triangles), S2 (yellow diamonds), L1.8 (purple stars), L2 (light blue pentagons), T2 (dark blue squares), SB-AGN (green circles), and SB (green circles with small black dots).} \label{fig:MeanANN3vsLhard} \end{figure} One of the main differences between the X-ray spectra of the various optical classes is set by the $\rm{\nu_{SB}}$ component, which is increasing from the S1 to the SB classes, passing through the S1.8, S2, L1.8, and L2/T2/SB-AGN groups. The nature of the $\rm{\nu_{SB}}$ component cannot be fully assessed with the results of this analysis alone, but we discuss below possible explanations. The star-formation (circumnuclear or that of the host galaxy) is the most natural explanation for the $\rm{\nu_{SB}}$ component. In this case, X-ray emission by binary systems, supernovae remnants, and/or emission by diffuse hot gas, could contribute to this $\rm{\nu_{SB}}$ component. In this case we would expect $\rm{\overline{\nu}_{SB}}$ to increase when the luminosity decreases for the objects in our sample. To test such hypothesis Fig. \ref{fig:MeanANN3vsLhard} shows the average $\rm{\nu_{SB}}$ ($\rm{\overline{\nu}_{SB}}$) versus the mean value for log(L(2-10 keV)). These two quantities are in fact clearly anti-correlated (r=0.94, $\rm{P_{null}=5\times10^{-5}}$)\footnote{Note that $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ show a poor relation with $\rm{log(\overline{L}(2-10 keV))}$ in the top panel of Fig. \ref{fig:MeanANN3vsLsteep}.}. Thus, $\rm{\nu_{SB}}$ increases when the X-ray luminosity decreases, in favour of our hypothesis that the $\rm{\nu_{SB}}$ component is related to star-formation. The SB galaxies, with the largest $\rm{\nu_{SB}}$ values in our sample, have an X-ray luminosities of $\rm{\sim 10^{40} erg~s^{-1}}$. This could be representative of the galactic X-ray emission due to the processes mentioned above. If an AGN component is present in almost all galaxies, then as it becomes stronger, $\rm{\nu_{SB}}$ decreases, while at the same time the X-ray luminosity increases. The $\rm{\nu_{SB}}$ component is almost zero in the S1.8 and S1 classes probably because the AGN-like source entirely outshines the underlying host-galaxy emission, or it could also mean that the $\rm{\nu_{SB}}$ component is entirely absent. For example, \citet{Wu09} (and references therein) claimed that the circumnuclear star-formation might be even destroyed in the presence of an AGN. An alternative origin for the $\rm{\nu_{SB}}$ component for those sources hosting an AGN is the X-ray emission from the hot plasma in the NLR, emission from the ``scattering component'' in AGN or ionised gas. It has been claimed that high resolution X-ray spectra are dominated by emission lines from the NLR in Type-2 Seyferts \citep{Guainazzi07}. Moreover, the soft X-ray emission in a few AGN is extended on scales ranging from a few hundred parsecs to a few thousand parsecs, in close agreement with the morphology of the NLR seen at optical wavelengths for both LINERs and Type-2 Seyferts \citep[][]{Gonzalez-Martin10,Bianchi06,Masegosa11}. In this case we would expect $\rm{\overline{\nu}_{SB}}$ to increase when the luminosity increases for the objects in our sample. However, as mentioned before, $\rm{\nu_{SB}}$ increases when the X-ray luminosity decreases (see Fig. \ref{fig:MeanANN3vsLhard}), ruling out the NLR as the main responsible for the $\rm{\nu_{SB}}$ component. \subsection{Elusive AGN}\label{sec:hidden} The $\rm{\nu_{S1}}$ and/or $\rm{\nu_{S1.8}}$ components are not negligible in most of the emission line nuclei presented in this paper (see Figs. \ref{fig:ANNhist} and \ref{fig:ANNnuvsnu}). A total of 22 out of the 162 spectra (i.e., 13.5\%) are consistent with no signature of an AGN-like component; this percentage is slightly higher in terms of the number of objects (19 out of the 90, 21\%). Thus, $\rm{\sim}$80\% of our sample show signs of an AGN-like component, either with a S1-like or a S1.8-like contribution. This number is almost twice the percentage of AGN (43\%) estimated at optical frequencies by \citet{Ho97} for the same sample. Moreover, although for some T2, SB-AGN, and SB nuclei, the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ components are consistent with zero, on average, the X-ray spectra of these nuclei do show the presence of $\rm{\nu_{S1}}$ or $\rm{\nu_{S1.8}}$ components. However, based on their optical spectra, these classes correspond, at best, to objects on the border between AGN and star-forming galaxies. Our result strongly supports the hypothesis that an AGN component might be present at X-rays at a certain level in most of the emission line nuclei included in our sample, even if they do not show signatures of this AGN component in their optical spectra. Non-AGN at optical wavelengths with AGN signatures at X-rays have been largely studied in the literature \citep[called `elusive AGN', see][]{Maiolino98,Soria06A,Soria06B}. Galaxies with bulges harbour BHs \citep[see][and references therein]{Kormendy13}. However, at optical wavelengths, only a small fraction of bulge galaxies show evidence for AGN activity; in about half of the high signal-to-noise (S/N) ratio optical spectra taken by \citet{Ho97} there is no indication of AGN activity. \citet{Tzanavaris07} studied a sample of star-forming galaxies classified by \citet{Ho97} at X-ray, finding AGN signatures for a large fraction of them. This is consistent with our results. \citet{Tzanavaris07} suggested that the lack of optical signatures may be due to the fact that the emission could be overwhelmed by that coming from circumnuclear star formation. This is entirely consistent with the increase of the $\rm{\nu_{SB}}$ component when the luminosity decreases (see Fig. \ref{fig:MeanANN3vsLhard} and previous Section), if the $\rm{\nu_{SB}}$ component is associated with the constant, diffuse X-ray emission of the host galaxy and/or X-ray emission associated with intense star-forming regions. \subsection{Relevance of the ANN method for X-ray surveys}\label{sec:surveys} ANNs have proven to be a powerful approach to a broad variety of problems \citep[e.g.,][]{Bishop96, Gupta04, Asensio-Ramos05, Socas-Navarro05, Carballo08, Han12}. In the most common application, ANN functions as a classification algorithm. In the AGN field, for instance, \citet{Rawson96} already used the ANN to classify optical spectra into Type-1 and Type-2 AGN. However, ANN have not been used to classify X-ray spectra before. Using other statistical methods, several attempts have been made to classify X-ray spectra, particularly for low S/N spectra. \citet{Norman04} selected normal, Type-1 and Type-2 AGN galaxies from the \emph{Chandra} Deep field North (CDF-N) and South (CDF-S) samples using a Bayesian classification procedure. Priors were constructed from a set of galaxies with well-defined optical classes. They used the X-ray hardness ratio, the 0.5-2 keV X-ray luminosity, and the ratio between X-ray and optical fluxes. The product of the prior distribution for a class and the likelihood for the observed parameters for a given source gave the probability that the source was drawn from that class. \citet{Ptak07} used a similar methodology with several improvements (e.g., \emph{k}-correction in the optical data). They showed that the method was efficient in classifying the X-ray spectra into Type-1, Type-2 and normal galaxies. Our methodology has two advantages: (1) it does not need any optical information and (2) it is able to distinguish among the S1, S1.8, L1.8, S2, L2/T2/SB-AGN, and SB classes. We show that the ANN is an excellent tool to discriminate between most of the optical classes using only their X-ray spectra. It might be very useful for X-ray surveys where the optical information is missed. The ANN components can be computed for any set of X-ray spectra using our already trained ANN\footnote{We kindly suggest to contact any of the coauthors of the paper for the use of our trained ANN.}. The effects of using X-ray spectra with lower S/N to their classification with the ANN method needs to be explored (perhaps through simulations), which is out of the scope of this paper. Finally, the ANN should be able to classify objects in broad classes, and the results will be useful for statistical studies. However, the method is not being demonstrated to be particularly useful in the classification of objects on an individual basis. We have investigated the connection between optical classes and X-ray spectra in a sample of 90 nearby emission line galaxies. We have used flux-calibrated X-ray spectra observed with \emph{XMM}-Newton/pn. The results of this paper are, for the first time, free of the subjectivity of the X-ray spectral fitting thanks to the use of the ANNs: \begin{itemize} \item We used a set of the S1, S1.8 and SB classes to train the ANN, giving as output arrays $\rm{\nu_{S1}}$, $\rm{\nu_{S1.8}}$, and $\rm{\nu_{SB}}$, respectively. The ANN is 90\% efficient to distinguish these classes. They all show then distinctive signatures at X-rays. \item Based on their X-ray spectral shape, the emission line nuclei in the nearby galaxies are divided into six groups: S1, S1.8, S2, L1.8, L2/T2/SB-AGN, and SB classes. Only the L2, T2, and SB-AGN classes show the same average X-ray spectrum even though they belong to distinct optical classes. Furthermore, the objects within each of these six classes have similar average X-ray spectra. \item The average X-ray spectrum of the objects in each X-ray class can be described by the contribution of two components, either the $\rm{\nu_{SB}}$ and $\rm{\nu_{S1}}$, the $\rm{\nu_{SB}}$ and $\rm{\nu_{S1.8}}$, or the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ (in the case of S1s and S1.8s). The S2 (L1.8) class is similar to the S1.8 (S1) class but with larger contributions of the $\rm{\nu_{SB}}$ component. The L2/T2 and SB-AGN classes have a strong $\rm{\nu_{SB}}$ component, with the addition of a $\rm{\nu_{S1.8}}$ component. \item The S1 and S1.8 classes show little $\rm{\nu_{SB}}$ and a wide range of the $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ components. We show that this wide range of $\rm{\nu_{S1}}$ and $\rm{\nu_{S1.8}}$ contributions is most probably related to the different amount of obscuration that affects the nuclear emission at X-rays, in agreement with the UM predictions. \item Most of the objects in our sample have a non negligible contribution of either a $\rm{\nu_{S1}}$ or a $\rm{\nu_{S1.8}}$ component. This result strongly supports the presence of an AGN-like nucleus in most nearby galaxies, albeit at different levels of luminosities (i.e. activity). \item We argue that the $\rm{\nu_{SB}}$ component is associated to a contribution of star-formation in the host-galaxy. As the contribution of the AGN component decreases, the $\rm{\nu_{SB}}$ component increases, and at optical wavelengths it shows stronger signatures representative of S2, L1.8, L2/T2/SB-AGN, and finally of SB nuclei. \end{itemize} We find that the emission line nuclei in nearby galaxies can be classified in six classes, based on the shape of their X-ray spectra. These classes are associated to the traditional optical classes, although their number is smaller. Thus, the shape of the X-ray spectra of those galaxies may be determined by a smaller number of physical parameters than those which determine the optical classes. Alternatively, this could be due to the difficulties to classify them at optical wavelength using the BPT diagrams. Indeed, our results suggest that the X-ray spectra of nearby galaxies are simply the combination of two components. The first one is an AGN-like component and the second one is due to star-formation in the host-galaxy contributing to the X-rays. An AGN-like nucleus may be present in most of them (80\%). Its strength, relative to the contribution of star-formation in the host-galaxy, determines the average X-ray spectrum of objects for each X-ray class. A third physical parameter could be related to the amount of obscuring material along the LOS. This parameter almost certainly drives the Type-1/Type-2 dichotomy, but may also explain why, for example, the L1.8 class predominantly shows a $\rm{\nu_{S1}}$ component in their spectra while L2, T2, and SB-AGN predominantly show a $\rm{\nu_{S1.8}}$ component. We conclude that the ANN method is quite powerful to detect AGN-like nuclei (and distinguish which ones are affected by absorption). It can therefore be used to identify AGN, and even to infer their optical classes, using only X-ray spectra and our trained ANN. However, this can only be done in a statistical way, i.e., using the X-ray spectra of a large number of objects. This methodology could be very useful in X-ray surveys, e.g. the \emph{eRosita} survey, where the optical information for tens of thousands newly discovered objects will not be available. | 14 | 4 | 1404.1078 |
1404 | 1404.6144_arXiv.txt | {We present an analysis of the debris disc around the nearby K2~V star HIP~17439. In the context of the \emph{Herschel}\thanks{\emph{Herschel} is an ESA space observatory with science instruments provided by European-led Principal Investigator consortia and with important participation from NASA.} DUNES key programme the disc was observed and spatially resolved in the far-IR with the \emph{Herschel} PACS and SPIRE instruments. In a first model, Ertel et al. (2014) assumed the size and radial distribution of the circumstellar dust to be independent power laws. There, by exploring a very broad range of possible model parameters several scenarios capable of explaining the observations were suggested. In this paper, we perform a follow-up in-depth collisional modelling of these scenarios trying to further distinguish between them. In our models we consider collisions, direct radiation pressure, and drag forces, i.e. the actual physical processes operating in debris discs. We find that all scenarios discussed in Ertel et al. are physically sensible and can reproduce the observed spectral energy distribution along with the PACS surface brightness profiles reasonably well. In one model, the dust is produced beyond $120\au$ in a narrow planetesimal belt and is transported inwards by Poynting-Robertson and stellar wind drag. A good agreement with the observed radial profiles would require stellar winds by about an order of magnitude stronger than the solar value, which is not supported~-- although not ruled out~-- by observations. Another model consists of two spatially separated planetesimal belts, a warm inner and a cold outer one. This scenario would probably imply the presence of planets clearing the gap between the two components. Finally, we show qualitatively that the observations can be explained by assuming the dust is produced in a single, but broad planetesimal disc with a surface density of solids rising outwards, as expected for an extended disc that experiences a natural inside-out collisional depletion. Prospects of discriminating between the competing scenarios by future observations are discussed. | HIP~17439 (HD~23484) is a nearby solar-type star with an estimated age of $0.8$--$3.7\,\text{Gyr}$ \citep{Mamajek2008,Garces2010,Fernandes2011}. The first clear evidence for a circumstellar dust disc around HIP~17439 was found by \cite{Koerner2010} who detected a far-IR excess in \emph{Spitzer}/MIPS data. Recently, the HIP~17439 system was observed as part of the \emph{Herschel} \citep{Pilbratt2010} Open Time Key Programme DUst around NEarby Stars \citep[DUNES,][]{Eiroa2010,Eiroa2013}. The disc appeared spatially resolved in the far-IR images of the PACS \citep{Poglitsch2010} and SPIRE \citep{Griffin2010,Swinyard2010} instruments. Together with the discs around HD~202628 \citep[][]{Krist2012}, HD~207129 \citep{Krist2010,Marshall2011,Loehne2012}, and HD~107146 \citep{Ertel2011}, HIP~17439's disc is amongst the most extended around sun-like stars identified to date. Theory predicts that the dust in such a disc is short-lived compared to the lifetime of the star, owing to the mutual collisions and radiative forces. Therefore, the dust is thought to be continually replenished through collisional grinding of km-sized asteroidal bodies (planetesimals) that are remnants of the planet formation process \citep{Backman1993, Wyatt2008, Krivov2010}. In the DUNES programme, those debris discs have been detected with a rate of at least 20\% around solar-type main-sequence stars \citep{Eiroa2013}. Their presence may or may not be correlated in one or another way with the presence of planets \citep{Moro-Martin2007, Kospal2009, Bryden2009,% Maldonado2012, Wyatt2012,Matthews2014,Marshall2014}. The modelling of the spectral energy distribution (SED) of a debris disc is in general a degenerate problem. For example, the observed flux can be reproduced by the thermal emission of large dust grains close to the star or small ones located far away. A spatially resolved debris disc where the location of the dust is directly measurable offers the best opportunity to break those degeneracies. Modelling of resolved systems places meaningful constraints on the discs' radial and temperature structures as well as the grain sizes and compositions \citep[e.g.,][]{Matthews2010,Ertel2011,Eiroa2011,Lestrade2012,Booth2013}. Simultaneous modelling of the SED and the radial surface brightness profiles extracted from the PACS images of the HIP~17439 system was done by \cite{Ertel2014} for the first time. There, a dust surface number density $n\propto s^{\gamma} r^{\alpha}$ was assumed, i.e. a combination of two independent power laws for the size distribution (exponent $\gamma$) and the radial distribution (exponent $\alpha$) of the grains, fitted to the observational data by a multi-wavelength $\chi^2$ minimization. This way, one can readily explore a huge parameter space and thereby find the most appropriate disc configurations. \citeauthor{Ertel2014} showed that the SED and the radial profiles can be well reproduced by either a one- or a two-component scenario (Table~\ref{tab:steves_bestfit}). \vspace{-0.25cm} \begin{table}[htb!] \centering \caption{Best-fit results with $3\sigma$ uncertainties of the power-law models from \cite{Ertel2014}. The dust surface number density is defined as $n\propto s^{\gamma} r^{\alpha} $. $r_1$ and $r_2$ are the inner and outer boundary of a disc, $\theta$ is the disc inclination from face-on, and $M_\text{d}$ the dust mass for particles with radii $s<1\milli\metre$.} \renewcommand{\arraystretch}{1.5} \begin{tabular}{cccc} \hline\hline Parameter & \multicolumn{3}{c}{Best-fit} \\ \hline &\multicolumn{1}{c}{One-component model} & \multicolumn{2}{c}{Two-component model} \\ & & Inner disc & Outer disc \\ \hline $r_1$ [au] & \multicolumn{1}{c}{8.3$^{+5.6}_{-0.8}$ } & 29.2$^{+6.6}_{-27.5}$ & 90.9$^{+79.9}_{-74.9}$\\ $r_2$ [au] & \multicolumn{1}{c}{394.0$^{+106.0}_{-267.4}$} & 500.0$^\ast$ & 500.0$^\ast$ \\ $\alpha$ & \multicolumn{1}{c}{-0.1$^{+1.0}_{-1.5}$} & -4.0$^{+3.6}_{-1.0}$ & -1.6$^{+2.6}_{-3.4}$ \\ $s_\text{min}$ [$\!\microns$] & \multicolumn{1}{c}{8.1$^{+2.6}_{-1.9}$} & 5.2$^{+10.8}_{-1.7}$ & 12.4$^{+17.5}_{-12.3}$ \\ $\gamma$ & \multicolumn{1}{c}{-4.0$^{+1.0}_{-0.9}$} & -5.5$^{+1.8}_{-0.0}$ & -4.3$^{+1.3}_{-1.2}$\\ $\theta$ [deg] & \multicolumn{1}{c}{63.9$^{+18.1}_{-46.1}$} & \multicolumn{2}{c}{60$^{+10}_{-10}$}\\ $M_\text{d}$ $[M_\oplus]$ & \multicolumn{1}{c}{$1.3\times10^{-2}$} & $2.1\times10^{-4}$ & $1.1\times10^{-2}$ \\ Dust material & astrosilicate & \multicolumn{2}{c}{astrosilicate} \\ \hline \end{tabular} \label{tab:steves_bestfit} \medskip \tablefoot{Values marked with $^\ast$ were fixed in the modelling and are not outcomes of the fitting procedure. } \end{table} \vspace{-0.25cm} The best-fit one-component model found is a broad dust disc with a radial extension of several hundreds of au where the number density is nearly constant ($\alpha=-0.1$). Two possible morphologies of such a disc are conceivable. The first one is a narrow planetesimal belt near the outer disc edge. Through transport processes such as Poynting-Robertson (P-R) drag dust particles move inwards, filling the inner disc region and ensuring a broad distribution of dust. The second possibility is a broad planetesimal ring. This implies that dust is produced everywhere in the whole disc, which~-- under certain assumptions~-- may result in a constant dust surface density as well. The two-component model consists of two rings with inner edges around 29 and 91~au. Due to the large negative $\alpha$ values, most of the dust is located near the inner disc edges of the two components and the rings are spatially separated by a wide gap. Continuing the study by \cite{Ertel2014}, this paper presents an in-depth collisional modelling of the HIP~17439 system. To this end, we start with an initial distribution of planetesimals, which we also refer to as parent bodies, and consider their subsequent collisional evolution. As a result, the bodies are ground down to dust where the smallest particles end up with typical sizes in the order of $1\microns$ or smaller. The collisional outcomes depend on the size of the impactors, the material properties, and the relative velocities in the disc. Our modelling also includes dust transport mechanisms in the form of stellar wind and P-R drag, dependent on grain size, material properties and stellar distance. Hence, the dust radial and size distributions are intrinsically coupled. Their slopes can no longer be controlled directly, but instead, they are determined by the evolution of the planetesimal disc. This way, we model the actual physical processes operating in a debris disc. However, due to the numerical complexity of this method, we can only explore a limited number of parameter combinations. As a starting point, we use the one- and two-belt models from \cite{Ertel2014} and check whether these are physically plausible. In Sect.~\ref{sec:model_setup}, we describe the data used and the technique of our modelling. Sections~\ref{sec:onering} and \ref{sec:tworings} present simulations with one narrow parent belt and two such belts, respectively. Section~\ref{sec:extended_belt} discusses the possibility of a single extended planetesimal disc. Some prospects for future observations are given in Sect.~\ref{sec:prospects}. Section~\ref{sec:conclusion} contains conclusions and a discussion of our results. | \label{sec:conclusion} Our collisional modelling of the HIP~17439's debris disc does not allow us to draw any strict conclusions as to the underlying architecture of the system. The data can be fitted well with a single disc or two separated discs, with the radial distribution of the planetesimals being narrow in both cases. Furthermore, we argue that an extended planetesimal belt, approximated by a multicomponent disc, would reproduce the data as well. Below we discuss the astrophysical plausibility of all three models. \begin{enumerate} \item {\em One parent belt}. Provided that a planetesimal belt is located at a distance $>\!120\au$ from the star, suggested by the one-component model of \cite{Ertel2014}, dust material has to be transported efficiently from the planetesimal location to the inner region of the system, otherwise the inner disc's surface brightness is too shallow and contradicts the observed radial profiles. Since Poynting-Robertson drag alone is not sufficient enough, we found a need for strong stellar winds. We identified a wind strength of 15 times the solar value as the upper limit to prevent the radial profiles from being steepened too much if the inner edge of the planetesimal belt lies closer than $150\au$ to the star. The question remains whether HIP~17439 does possess winds of that strength. In our best run the planetesimals reside between 150 and 180~au, i.e. they are much closer to the star than in the one-component scenario of \cite{Ertel2014}. There, the outer ring edge was found at about 400~au, but could also be located much further inside due to large uncertainties. In general, planetesimals and dust production at very large stellar distances seem to be unlikely because of the increasingly long growth and stirring timescales of planetesimals \citep[e.g.,][]{Kenyon2008}. Hence, the moderate radial range of planetesimals in the one-belt model presented in this study better fits in the debris formation theory. \item {\em Two parent belts}. Without stellar winds of sufficient strength, the data are consistent with a two-ring disc, with a warm inner and a cold outer component. Many systems are believed to have a two-belt configuration since their SEDs can be well fitted by using two blackbody curves \citep[e.g.,][]{Matthews2010,Donaldson2013,Broekhoven-Fiene2013,Su2013}. Surveys of two component discs around stars of different spectral types highlight that inner and outer component have median temperature values of $\approx\!190~\kelvin$ and $\approx\!60~\kelvin$, respectively \citep[e.g.,][]{Morales2011,Ballering2013}. This yields a distance ratio of $r_\text{outer}/r_\text{inner}=(190/60)^2\approx10$, indicating a distinct gap between both components. One possible explanation of this result considers planets which have formed within the discs and split them up \citep[e.g.,][]{Ertel&Wolf&Rodmann2012}. In our best-fit two-belt model (model~II) the mean planetesimal distances for the inner and outer disc have a ratio of $\approx\!5$, i.e. close to what was found for many other systems. By testing different maximum eccentricities $e_\text{max}$ of the inner and outer planetesimal belt, we discovered that both components must have similar dynamical excitation of $e_\text{max}\approx0.04$ to be in good agreement with the radial profile data. One problem is that this level, especially in the inner belt, is lower than what is expected for the stirring by a planet \citep[e.g.,][]{Mustill2009}, suggesting that the gap between the two belts may be not populated by planets. In that case, however, it would be difficult to explain what else, if not planets, has cleared up the wide gap between the two belts. Nevertheless, the problem can be mitigated by the assumption that possible planets in the gap are in nearly circular orbits and/or have low masses. \item {\em One extended parent belt}. We discuss this possible disc architecture by using more than two belts adjacent to each other as a proxy. In principle, this could be the best model consistent with planet(esimal) formation theories. The optical thickness profile would be nearly constant over a wide radial range, starting to decrease beyond $\approx\!200\au$. The radial surface density profile of the underlying planetesimal disc is rising outwards, consistent with a long-term inside-out collisional erosion of such a disc. Therefore, from the point of view of collisional modelling we can also confirm the extended planetesimal belt hypothesis already proposed in \cite{Ertel2014}. \end{enumerate} More observations are required to discriminate between the competing scenarios discussed in this paper. LMT and ALMA are the most promising facilities to shed light on the actual structure of HIP~17439's debris disc. We simulated LMT and ALMA images of our best one- and two-belt models at millimetre/submillimetre wavelengths. These tests highlight that present-day telescopes are possibly capable to distinguish between a one- or a two-belt model but only with high observational effort. \begin{acknowledgement} We thank the reviewer for a speedy and constructive report that helped to improve the manuscript. CS, TL, and AVK acknowledge support by the \emph{Deut\-sche For\-schungs\-ge\-mein\-schaft} (DFG) through projects Kr~2164/10-1 and \mbox{Lo~1715/1-1}. SE thanks the French National Research Agency (ANR, contract ANR-2010 BLAN-0505-01, EXOZODI) and PNP-CNES for financial support. JPM and CE are partly supported by Spanish grant AYA 2011-26202. \end{acknowledgement} | 14 | 4 | 1404.6144 |
1404 | 1404.1308_arXiv.txt | Recent simulations of the densest portion of the Corona Borealis supercluster (A2061, A2065, A2067, and A2089) have shown virtually no possibility of extended gravitationally bound structure without inter-cluster matter (Pearson \& Batuski). In contrast, recent analyses of the dynamics found that the clusters had significant peculiar velocities towards the supercluster centroid (Batiste \& Batuski). In this paper we present the results of a thorough investigation of the CSC: we determine redshifts and virial masses for all 8 clusters associated with the CSC; repeat the analysis of Batiste \& Batuski with the inclusion of A2056 and CL1529+29; estimate the mass of the supercluster by applying the virial theorem on the supercluster scale (e.g. Small et al.), the caustics method (e.g. Reisenegger et al.), and a new procedure using the spherical collapse model (SCM) with the results of the dynamical analysis (SCM+FP); and perform a series of simulations to assess the likelihood of the CSC being a gravitationally bound supercluster. We find that the mass of the CSC is between \mbox{$0.6$ and $12 \times 10^{16} \, h^{-1} \, \mathrm{M}_{\sun}$}. The dynamical analysis, caustics method and the SCM+FP indicate that the structure is collapsing, with the latter two both indicating a turn around radius of \mbox{$\sim 12.5 \, h^{-1} \, \mathrm{Mpc}$}. Lastly, the simulations show that with a reasonable amount of inter-cluster mass, there is likely extended bound structure in the CSC. Our results suggest that A2056, A2061, A2065, A2067, and A2089 form a gravitationally bound supercluster. | Superclusters of galaxies are the largest coherent structures in the Universe, and studies \citep{Rood76} have shown that their internal dynamics are generally dominated by Hubble flow (i.e. not gravitationally bound). Extended bound structure within superclusters is unusual, but intensive studies of the Shapley Supercluster (SSC) (e.g. \cite{Bardelli93,Proust06}, and references therein) have demonstrated significant bound structure, with a central core of five clusters that is in the final stages of collapse \citep{Reisenegger00,Munoz08,Pearson13}. To date the SSC is the only confirmed bound supercluster in the Universe (for the purposes of the current paper we define a supercluster as containing no fewer than $3$ rich clusters of galaxies), but its existence suggests that similar structures might be found elsewhere in the Universe. The Corona Borealis supercluster (CSC) is a particularly dense, compact supercluster that has been identified as a candidate bound supercluster similar to the SSC. It was included by \citet{Abell61} in his catalog of second order clusters and first noted by \citet{Shane59}, who initially identified 12 member clusters in a \mbox{$6\degr \times 6\degr$} region before later using brightest cluster galaxies to show that there were actually two components viewed in projection. The foreground component, at \mbox{$z\approx 0.07$}, contains Abell clusters 2056, 2061, 2065, 2067, 2079, 2089 and 2092 and is what we now refer to as the CSC. \citet{Postman88} performed the first dynamical analysis of this region, making virial mass estimates of six clusters (excluding A2056), finding a mass for the supercluster of \mbox{$8.2 \times 10^{15} \, h^{-1} \, \mathrm{M}_{\sun}$}. The Norris Survey \citep{small1,small2,small3} expanded on this work, performing N-body simulations to test mass estimators, demonstrating that the virial theorem was applicable on supercluster scales. By treating the six clusters in the core as a single virialized system, they estimated a mass of at least \mbox{$3 \times 10^{16} \, h^{-1} \, \mathrm{M}_{\sun}$}, concluding that the system was bound and had likely reached turnaround. \citet{Kopylova98} performed a dynamical analysis of the CSC using redshift independent distance estimates to assess peculiar velocities for eight clusters in the region (including A2019 and A2124 in addition to those of Postman et al.). Their results led them to define a rapidly collapsing core containing five clusters; A2061, A2065, A2067, A2089 and A2092. Recently \cite{Pearson13} (hereafter P13) and \cite{Batiste13} (hereafter B13) have revisited these results with the intention of more accurately assessing the current dynamical state of the CSC using currently available observational data. P13 performed N-body simulations of the CSC using the most accurate cluster mass estimates available for A2061, A2065, A2067 and A2089 (and assuming negligible inter-cluster mass), concluding that there is very little likelihood that any part of the structure is bound. B13 used Sloan Digital Sky Survey (SDSS) data \citep{SDSS7} to perform a dynamical analysis with the six clusters used by Postman et al., finding peculiar velocities indicative of extended bound structure. In this paper we aim to address the apparent conflict between these results, providing a more complete picture of the current dynamical state of this structure, and a reasonable estimate of the mass of the bound portion. We undertake an extensive analysis, employing several independent methods of mass estimation, and performing a large number of simulations with which we can assess the validity of the conclusions we draw from the dynamical analysis. The breadth of the analysis is intended to provide a context within which the affects of the assumptions underlying each method can be assessed, and the significant uncertainties on each result can be interpreted. The paper is structured as follows: In section 2 we describe the different data sets used for mass estimation and dynamical analysis, and explain the selection criteria and corrections to photometry. In Section 3 we present the dynamical analysis. In section 4 we present the mass estimation methods, which constrain the mass for several different scenarios: upper and lower bounds are placed on the possible mass of the CSC; the mass required to generate the motions reported in Section 3 is determined; and an alternative assessment of the mass is made using an independent set of observational data. In section 5 we present the results of our simulations. We use these results to assess the likelihood that the CSC is a bound structure, and to investigate whether inter-cluster dark matter would need to contribute significantly to the total supercluster mass. Lastly, in section 6 we discuss the implications of our results. Throughout this paper we adopt a standard \mbox{$\Lambda\mathrm{CDM}$} cosmology with \mbox{$H_{0}=100 \, h \, \mathrm{km} \; \mathrm{s}^{-1} \, \mathrm{Mpc}^{-1}$}, \mbox{$h = 0.7$}, \mbox{$\Omega_{\Lambda,0} = 0.7$}, \mbox{$\Omega_{m,0} = 0.3$} and \mbox{$q_{0}=-0.55$}. | It is fairly straightforward to place firm limits on the possible mass of the CSC. Application of the virial theorem sets an upper bound of \mbox{$1.2 \times 10^{17} \, h^{-1} \, \mathrm{M}_{\sun}$}, and summing the cluster masses (excluding CL1529+29) gives a lower bound of \mbox{$5.9 \times 10^{15} \, h^{-1} \, \mathrm{M}_{\sun}$}. These bounds cover a significant range, and it is unlikely that either method provides an accurate assessment of the mass. Given the probability that inter-cluster mass contributes significantly to the total mass of the CSC \citep{Proust06}, it is all but certain that the mass is larger than that obtained by summing the cluster masses. Similarly, the virial theorem assumes a bound virialised structure extending over a significant range in velocity space, and therefore very likely overestimates the total mass of the CSC. The dynamical analysis and resulting simulations and mass estimation, as well as the independent method of mass estimation provided by the caustics method, are intended to provide a more physically meaningful assessment of its mass and bound extent. The FP analysis indicates significant extended bound structure in the CSC, suggesting two bound regions that have reached turnaround and are in collapse: the first consisting of A2065, 2056 and 2089, and the second consisting of 2061 and 2067, and possibly also 2092. There is also some indication that these regions may not be dynamically isolated, and the CSC may consist of a single collapsing core of at least five clusters. Simulations informed by the FP results, and including no inter-cluster mass, support these conclusions, indicating a significant likelihood \mbox{($> 50$ per cent)} of each of the two cores being gravitationally bound, and some likelihood \mbox{($\sim 18$ per cent)} of the two regions being bound to each other. The errors on the FP distance estimates are quite low and appear to be free of any systematic bias, and the resulting cluster peculiar velocities are consistent with what might be expected for a bound structure of this type (see discussion in B13). However it is important to consider that the distance errors are still too large compared to the peculiar velocities to give a definitive analysis of the dynamics, so the results from extensive simulations that account for these errors are key in assessing the validity of the FP results and the conclusions we draw from them. The two most physically meaningful methods of mass estimation are the SCM+FP, which is based on the FP results but also accounts for the errors in the distance determinations, and the caustics method, which provides an independent method of mass estimation based on observational data. The SCM+FP analysis indicates a bound region with a mass of \mbox{$0.91 \times 10^{16} \, h^{-1} \, \mathrm{M}_{\sun}$} within a turn-around radius of \mbox{$12.5 \, h^{-1} \, \mathrm{Mpc}$}, while the caustics method indicates a bound region with a mass of \mbox{$1.02 \times 10^{16} \, h^{-1} \, \mathrm{M}_{\sun}$} within a turn-around radius of \mbox{$12.6 \, h^{-1} \, \mathrm{Mpc}$}. The remarkable agreement between the mass estimates and turn-around radii from these methods is particularly significant. Given the fundamental differences between the data sources and methods of analysis, it is highly unlikely that this consistency is the result of some underlying systematic bias. Consequently, each of these methods can be viewed as independently verifying the other. Since the SCM+FP method is informed by the FP, this consistency suggests that our method of calibration is valid, and that the FP analysis accurately assesses the internal dynamics of the CSC. Taken together, these results indicate that the CSC contains a single bound core containing: A2056, A2061, A2067, A2089, and A2065, which has reached turnaround and is in collapse with a mass of \mbox{$\sim 1 \times 10^{16} \, h^{-1} \, \mathrm{M}_{\sun}$}. The results of Section \ref{sims:ICM} demonstrate that reasonable assumptions about inter-cluster mass result in a significant likelihood of extended bound structure in the CSC. The low mass halo simulations assume an inter-cluster matter component that requires no dark matter, and results in a \mbox{$\sim 60$ per cent} chance of A2056, A2065, A2067 and A2089 being bound, with a significantly lower likelihood that A2061 is part of the structure. Given the large tangential peculiar velocity of A2061 that results from the FP analysis, and its proximity to A2067, the simulations probably under-estimate the likelihood that they are a bound pair. The higher mass halo simulations would not be unreasonable assuming the presence of inter-cluster dark matter, and in that case the likelihood of the aforementioned clusters forming a bound structure is \mbox{$\sim 95$ per cent}, with A2061 showing a lower probability for the reasons given above. We also note that, if there is this much inter-cluster mass present, there is a good chance \mbox{($\sim 68$ per cent)} that A2092 is also part of the structure. Combining the results of the FP analysis, caustics method, SCM+FP, and simulations, we conclude that A2056, A2061, A2065, A2067, and A2089 comprise a gravitationally bound supercluster core. Should there exist an inter-cluster dark matter component, A2092 may also be part of this structure. This work provides the most conclusive evidence to date that the CSC is a bound supercluster similar to the SSC, and suggests that such structures may be found elsewhere in the Universe. | 14 | 4 | 1404.1308 |
1404 | 1404.6002_arXiv.txt | The origin of bipolar outflow asymmetry in young stellar objects (YSOs) remains poorly understood. It may be due to an intrinsically asymmetric outflow launch mechanism, or it may be caused by the effects of the ambient medium surrounding the YSO. Answering this question is an important step in understanding outflow launching. We have investigated the bipolar outflows driven by the T Tauri star DG Tauri on scales of hundreds of AU, using the Near-infrared Integral Field Spectrograph (NIFS) on Gemini North. The approaching outflow consists of a well-collimated jet, nested within a lower-velocity disc wind. The receding outflow is composed of a single-component bubble-like structure. We analyse the kinematics of the receding outflow using kinetic models, and determine that it is a quasi-stationary bubble with an expanding internal velocity field. We propose that this bubble forms because the receding counterjet from DG Tau is obstructed by a clumpy ambient medium above the circumstellar disc surface, based on similarities between this structure and those found in the modeling of active galactic nuclei outflows. We find evidence of interaction between the obscured counterjet and clumpy ambient material, which we attribute to the large molecular envelope around the DG Tau system. An analytical model of a momentum-driven bubble is shown to be consistent with our interpretation. We conclude that the bipolar outflow from DG Tau is intrinsically symmetric, and the observed asymmetries are due to environmental effects. This mechanism can potentially be used to explain the observed bipolar asymmetries in other YSO outflows. | \label{sec:intro} Outflows are ubiquitous components of young stellar objects (YSOs). Solar-mass YSOs are capable of driving collimated bipolar outflows of atomic and molecular material to distances of $\sim 1\textrm{ pc}$ \citep[e.g.,][]{MRF07}. These outflows extract angular momentum from the star-disc system, allowing material from the circumstellar disc to accrete onto the central protostar. It is generally accepted that the outflows are launched magnetocentrifugally, either from the surface of the circumstellar disc \citep[MHD disc wind;][]{BP82,PN83}, or from reconnection points in the stellar magnetosphere \citep[the X-wind;][]{Se94}. Multiple launch mechanisms may act in concert to produce outflows with multiple velocity components \citep{Ae03,L03,FDC06,SLH07}. The launch region of these outflows is unresolvable with current telescopes. Therefore, detailed observational and theoretical studies of YSO outflows are necessary in order to determine the manner in which they are launched, and the physical conditions at their launching point(s). Bipolar outflow asymmetry is a common occurrence in large-scale YSO outflows. These outflows, which are characterised by the presence of shock-excited Herbig--Haro (HH) objects, are often observed to be one-sided, with only a blueshifted, or approaching, outflow visible \citep[e.g.,][]{EM98,MR04,MRF07}. Such an asymmetry may be caused by the circumstellar disc obscuring the receding component of a symmetric bipolar outflow. Alternatively, the receding outflow may have entered the dense molecular cloud complex behind the YSO, obscuring it from observation \citep{MRF07}. However, radial velocity asymmetry is often seen in objects with observable bipolar HH outflows. For example, \citet{Hie94} found that, of 15 T Tauri stars with observed bipolar HH outflows, 8 showed bipolar velocity asymmetry between the blueshifted and redshifted outflows. The ratio of radial velocities between the opposing outflows in these objects is in the range 1.4--2.6. Further studies have found more asymmetric bipolar HH outflows, such as HH 30, which has a radial velocity ratio $\sim 2$ between the two outflows \citep{Ee12}. One-sided knot ejections have also been detected, such as that from the driving source of the HH 111 outflow \citep{GRL12}. Asymmetrical knot ejections and differing mass outflow rates between the two sides of the bipolar outflow from the Herbig Ae star HD 163296 have also been detected \citep{We06}. This evidence raises the question of whether the asymmetry is caused by environmental effects \citep{Hie94}, or is an intrinsic feature of the outflows, either due to disc conditions (such as warping) affecting the outflow launching \citep{GRL12}, or to other effects close to the launch point \citep{We06}. With the advent of space-based telescopes such as the Hubble Space Telescope (HST), and the development of ground-based adaptive-optics systems, the large-scale outflows can be traced back to within a few hundred AU of the central protostar, and are observed as well-collimated `microjets'. These small-scale outflows provide an excellent laboratory for testing outflow launch models, as the outflow has yet to propagate to a distance where it interacts with the large-scale molecular cloud complex \citep[e.g.,][]{MRF07}. Therefore, the search for bipolar outflow asymmetry in these microjets is important in determining if the asymmetry is an intrinsic property of the outflows on all scales. Velocity asymmetries were observed in the profiles of bipolar forbidden emission line (FEL) regions in several young stars \citep{HMS97}. The forbidden emission lines trace the presence of microjets \citep[e.g.,][]{BE99,Be02,Ce04,Ce07}. Further long-slit optical and near-IR spectroscopic observations have confirmed kinematic and/or physical bipolar asymmetries occur in the outflows of DG Tauri B \citep{Pe11} and FS Tauri B \citep{Lie12}. There are conflicting views on the cause of these asymmetries. \citet{Pe11} argue that the asymmetry in DG Tau B is due to an asymmetric ambient medium, based on the observation that only one side of the bipolar outflow is driving the ambient medium into a CO outflow. \citet{Lie12} argue for a bipolar outflow that is being driven at a different mass-loss rate on either side of the circumstellar disc, with the velocity difference between the two jets keeping a linear momentum balance, so there is no observable recoil. It is important to differentiate between the possible causes of bipolar outflow asymmetry in order to determine if it is an intrinsic or an environmental effect. HST and adaptive optics also permits imaging and spectroimaging studies of YSO microjets \citep[e.g.,][]{Ke93,Le97}. Such studies have shown the presence of structural differences in the bipolar small-scale outflows from YSOs. The blueshifted collimated jet from the YSO HL Tauri is spatially coincident with an approximately axisymmetric bubble-like structure \citep{Te07}. Bipolar asymmetries in jet collimation are observed in the YSOs RW Aurigae \citep{YMe09} and DG Tau B \citep{Pe11}. Such studies have shown that structural bipolar outflow asymmetries are common in YSOs on the microjet scale \citep{Pe11}. Another example of bipolar asymmetry in T Tauri star microjets is the transitional Class I/Class II YSO DG Tauri. One of the most actively-accreting T Tauri stars, DG Tau has been used as a laboratory in searches for jet rotation \citep{B02,Pe04,Ce07}, jet knot generation \citep{LFCD00,Re12} and links between jet and disc properties \citep{Tes02}. DG Tau drives the blueshifted HH 158 \citep{MF83} and HH 702 \citep{Se03,MR04} outflows, the latter extending to a distance of $\sim 0.5\textrm{ pc}$ from the protostar. There is no known large-scale redshifted HH outflow associated with DG Tau. A bipolar microjet-scale outflow is present, and exhibits velocity asymmetry between the approaching and receding flows \citep{He94,Le97}. This was originally detected through long-slit spectroscopy of optical forbidden emission lines from the outflows \citep{Hie94}. The receding outflow was first imaged by \citet{Le97} using spectroimaging of [O I] 6300 \AA\ emission. They determined a radial velocity ratio of 1.4 between the approaching and receding outflows. More recently, \citet{A-Ae11} and \citet{MCW13a} (hereafter referred to as Paper I) detected structural and kinematic differences in the microjet-scale approaching and receding outflows of DG Tau, using spectroimaging data of spatially extended [Fe II] 1.644 $\mu$m line emission. The approaching outflow shows the classical YSO microjet morphology of a central, well-collimated, high-velocity jet with deprojected velocity $\sim 215\textrm{--}315\kms$. The jet is dominated by both stationary and moving shock-excited `knots' of emission. This jet is `nested' within a region of lower-velocity emission, which may be excited by the formation of a turbulent entrainment layer around the jet \citep[Paper I;][]{Pe03b}. A wide-angle approaching molecular wind is observed in H$_2$ 1-0 S(1) 2.128 $\mu$m line emission \citep[Paper I;][]{Be08}, providing a supply of material for the jet to entrain. On the other hand, the redshifted outflow shows no evidence of any jet-like components, and instead forms a large bubble-like structure. This was interpreted by \citet{A-Ae11} as being the counterpart `magnetic bubble' \citep{Cie09} to a similar structure they claimed $\sim 1\farcs 2$ from the central star in the approaching outflow channel. However, analysis in Paper I showed that they appeared to be interpreting the low-velocity entrainment component in that region as part of the central jet. We concluded that an approaching bubble structure does not exist (Paper I). Therefore, the nature and cause of the bipolar outflow asymmetry in DG Tau remains an open question. We investigate the bipolar asymmetry in the microjet-scale DG Tau outflows below, and conclude that environmental effects hamper the propagation of one side of an approximately symmetric bipolar outflow. We proceed as follows. In \S\ref{sec:obs}, we outline our multi-epoch NIFS observations and our data reduction procedure. \S\ref{sec:analysis} details our methods of analysing the data. In \S\ref{sec:bubble}, we argue that the structure is a stationary bubble with an internal velocity field describing expansion of gas towards the bubble walls, based on comparisons of the observed velocity structure to kinetic models. \S\ref{sec:bubbleAGN} outlines the results of simulations of bubbles driven by impeded active galactic nuclei (AGN) jets, and links this work to the morphology observed in the DG Tau receding outflow. We propose that the bubble in the DG Tau receding outflow is the result of a receding counterjet being obstructed by clumpy ambient material in the extended envelope around DG Tau \citep{KKS96a}. The receding outflow is currently in the momentum-driven bubble phase, similar to the simulations of radio galaxies by \citet{SB07} and \citet{WB11}. We construct an analytical model of an expanding jet momentum-driven bubble in \S\ref{sec:model}, and find that it predicts physical parameters consistent with those observed in the DG Tau bubble and the extended CO envelope around the system. Finally, in \S\ref{sec:D}, we discuss the impact of our results on the interpretation of bipolar outflow asymmetry in other YSOs, as well as the implications of episodic variability in YSOs on our model. We summarise our conclusions in \S\ref{sec:concl}. | \label{sec:concl} We have investigated the nature of the receding microjet-scale outflow of the YSO DG Tauri, utilising high-resolution spectroimaging data taken with the Near-infrared Integral Field Spectrograph (NIFS) on Gemini North. In the $H$-band, the outflow appears as a large bubble-like structure in [Fe II] 1.644 $\mu$m line emission. This is in stark contrast to the approaching outflow, which shows a two-component outflow, consisting of a high-velocity jet surrounded by a lower-velocity disc wind, stimulated into emission by turbulent entrainment (Paper I). In the $K$-band, `clumpy' H$_2$ 1-0 S(1) 2.1218 $\mu$m line emission is observed near the edge of the circumstellar disc, coincident with some of the brightest regions of redshifted [Fe II] emission. Line fits of this emission showed that the H$_2$-emitting gas is at the systemic velocity, indicating that it represents material stationary with respect to the central star instead of an outflow component. The emission-line velocity structure of the receding outflow is well-described by kinetic models of bubbles with an internal distribution of expanding, radiating gas. These models generate simulated emission line channel maps from input parameters such as bubble height, elongation, distance to the outflow source, and inclination of the outflow axis to the line of sight. Simulated IFU data produced by the models show our observational data are consistent with the presence of a stationary bubble, with an internal distribution of emitting gas expanding towards the bubble walls. We compared the current appearance of the receding DG Tau bubble with the four-stage evolutionary track for AGN jet-driven bubbles proposed by \citet{SB07}, and further refined by \citet{WB11}. We concluded that the DG Tau receding outflow is forming a momentum-driven bubble at the 2005 observing epoch, morphologically similar to the energy-driven bubble phase identified by \citet{SB07}. The receding counterjet is being blocked by a cloud of ambient medium in its path. As the jet begins to push past the clump, it blows a large bubble, which extends further than the distribution of material that is blocking its progress. We interpreted the observed H$_2$ emission as being indicative of such a clumpy medium above the surface of the circumstellar disc, which is interacting with both the jet and the expanding bubble. This interpretation is supported by the presence of an [Fe II] emission enhancement adjacent to one of the H$_2$ clumps, which suggests a jet-ambient medium interaction point. The presence of a clumpy, large-scale residual envelope of molecular material around DG Tau has been observed previously \citep{KKS96a}, lending further weight to this interpretation. Our multi-epoch data support this interpretation. Between 2005 and 2006, the bubble is effectively stationary, except for a small ($\sim 70\textrm{--}90\kms$) movement of the highest-velocity emitting material at the bubble apex. In 2009, this high-velocity material has disappeared. This may be indicative of the jet driving the bubble achieving breakout at some point after the 2006 observing epoch, leaving just the base of the bubble observable in the 2009 epoch. We constructed an analytical model of a jet momentum-driven bubble. This model describes the evolution of a bubble driven by a dispersed jet, propagating into a smooth ambient medium. Based on the physical scale of the receding DG Tau bubble, this model predicts that the ambient medium number density above the circumstellar disc surface is of order $10^6\textrm{ to }10^7\textrm{ cm}^{-3}$, which is in agreement with density estimates of the extended CO envelope around DG Tau. The model predicts a bubble expansion of $\lesssim 5\textrm{--}10\kms$ at the bubble head, which can be tested with multi-epoch observational data covering a five- to ten-year period. The cooling time of the bubble is estimated to be $\sim 26\textrm { yr}$. This is the time-scale on which the bubble disappears due to cooling once the jet head moves beyond the apex of the bubble. This also implies that the bubble is momentum-driven, as this cooling time is short compared to the estimated dynamical age of the bubble. Our conclusions do not necessarily exclude other proposed models for YSO bipolar outflow asymmetry. However, an asymmetric ambient medium obstructing the evolution of a symmetric bipolar outflow is the only model that explains the morphology of the receding DG Tau outflow. We have demonstrated that a long-slit spectrographic observation of DG Tau would weakly replicate the bipolar outflow velocity asymmetry observed in long-slit spectrographic observations of other YSOs, and is in agreement with the previous spectroscopic measurement for DG Tau \citep{Le97}. We have also shown that the outflow channel formed by an episode of outflow activity may close over during periods of relative outflow quiescence, based on observations of the large-scale DG Tau approaching outflow $\lesssim 0.5\textrm{ pc}$ from the central star. This implies that both the approaching and receding outflows from DG Tau may be forced to evolve through a bubble phase for any given outflow event. Multi-epoch data of DG Tau is the best method to test these conclusions. Our analysis has provided robust predictions of the evolutionary path that the bubble will take once the underlying jet breaks free of the ambient medium that obstructs its progress. These predictions may be directly compared to future observations. | 14 | 4 | 1404.6002 |
1404 | 1404.6528_arXiv.txt | \noindent We present simplified models for the galactic center $\gamma$-ray excess where Dirac dark matter annihilates into pairs or triplets of on-shell bosonic mediators to the Standard Model. These annihilation modes allow the dark matter mass to be heavier than those of conventional effective theories for the $\gamma$-ray excess. Because the annihilation rate is set by the dark matter--mediator coupling, the Standard Model coupling can be made parametrically small to `hide' the dark sector by suppressing direct detection and collider signals. We explore the viability of these models as a thermal relic and on the role of the mediators for controlling the $\gamma$-ray spectral shape. We comment on ultraviolet completions for these simplified models and novel options for Standard Model final states. | The particle nature of dark matter (\DM) remains one of the outstanding open questions in high energy physics. Experimental probes of the dynamics that connect the dark sector and the Standard Model (\SM) fall into three complimentary classes shown schematically in Fig.~\ref{fig:blob:diagrams} See \cite{Arrenberg:2013rzp} for a status report. Recent analyses of the \FERMI Space Telescope data find an excess of 1--10 \GeV $\gamma$-rays from the center of the galaxy. In fact, a similar excess seems to extend away from the center to high galactic latitudes~\cite{Hooper:2013rwa, Okada:2013bna, Huang:2013pda}. This may be indicative of dark matter annihilating into \SM final states which later shower to produce the observed excess photon spectrum \cite{Goodenough:2009gk, Hooper:2010mq, Abazajian:2010zy, Abazajian:2012pn, Boyarsky:2010dr, Hooper:2011ti, Gordon:2013vta, Macias:2013vya, Abazajian:2014fta, Daylan:2014rsa, Modak:2013jya}; see \cite{Anchordoqui:2013pta,Ko:2014gha,Buckley:2011mm,Boucenna:2011hy,Zhu:2011dz,Kyae:2013qna, Marshall:2011mm,Cerdeno:2014cda,Modak:2013jya, Agrawal:2014una} for recent models. While an early estimate argued that an alternate interpretation based on unidentified millisecond pulsars is unlikely \cite{Hooper:2013nhl}, \cite{Abazajian:2014fta} and \cite{Yuan:2014rca} recently demonstrated the consistency of this hypothesis with the $\gamma$-ray excess. Indeed, it may be difficult to distinguish these two possibilities since the extrapolated millisecond pulsar (\textsc{msp}) profile is very similar to standard \DM profiles \cite{UCI_MSP}. For the remainder of this paper we assume the excess is generated by \DM annihilation. The latest analyses prefer a 40~\GeV dark matter candidate that annihilates into $b\bar b$ pairs\footnote{% Annihilation of 10~\GeV \DM into $\tau\bar\tau$ is also plausible fit, see \cite{Hagiwara:2013qya,Buckley:2013sca,Buckley:2011mm,Boucenna:2011hy,Marshall:2011mm,Logan:2010nw} for recent models. \cite{Lacroix:2014eea} found that a universal coupling to charged leptons may be favored after bremsstrahlung and inverse Compton scattering effects are included. In this paper we focus on the case where the $\gamma$-ray excess is generated by $b\bar b$ pairs; we comment on more general final states in Section~\ref{sec:UV:MFV} and Appendix~\ref{app:spectral:shape}. } with a thermally averaged cross section $\langle \sigma v \rangle_{b\bar b} \approx \mathcal O(\text{few}) \times 10^{-26} \text{cm}^3/\text{s}$ \cite{Abazajian:2014fta, Daylan:2014rsa}. Further, because $\langle \sigma v \rangle_{b\bar b}$ is close to the value required to be a thermal relic from standard freeze-out, it is implausible that such a relic could produce such a $\gamma$-ray signal without having an $s$-wave annihilation mode. Combined with constraints from direct detection and collider experiments, this signal motivates a more detailed study of the physics encoded in the shaded regions of Fig.~\ref{fig:blob:diagrams}. \begin{figure}[t] \centering \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_blob_chichitoSMSM} \caption{} \end{subfigure} \qquad\quad \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_blob_chiSMtochiSM} \caption{} \end{subfigure} \qquad\quad \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_blob_SMSMtochichi} \caption{} \end{subfigure} \caption{(a) Annihilation, (b) Direct Detection, (c) Collider. Complimentary modes of dark matter detection. Annihilation sets both the thermal relic abundance and the present-day indirect detection rate.} \label{fig:blob:diagrams} \end{figure} \subsection{From Effective Theories to Simplified Models} A simple parameterization of the \SM--\DM interaction is to treat the shaded blobs as effective contact interactions between dark matter particles ($\chi$) and \SM states. For example, the coupling of fermionic \DM to a quark $q$ is parameterized through nonrenormalizable operators \begin{align} \mathcal L \supset \frac{1}{\Lambda^2} \left(\bar\chi \mathcal O_\chi \chi\right) \, \left(\bar q \mathcal O_q q\right), \label{eq:DM:EFT} \end{align} where, for example, $\mathcal O_\chi \otimes \mathcal O_q = \gamma^\mu \otimes \gamma_\mu$ corresponds to an interaction mediated by a heavy vector mediator that has been integrated out. The coefficient $\Lambda^{-2}$ can be calculated for specific \DM models and allow one to apply bounds from different types of experiments in a model-independent way. This technique has been applied, for example, for collider \cite{ Cao:2009uw, % Beltran:2010ww, % Goodman:2010yf, % Bai:2010hh, % Goodman:2010ku, % Fox:2011fx, % Rajaraman:2011wf, % Fox:2011pm, % Fortin:2011hv, % Bell:2012rg, % Cheung:2012gi, % Bai:2012xg, % Ding:2012sm, % Carpenter:2012rg, % Cotta:2012nj, % Zhou:2013fla, % Carpenter:2013xra, % Dreiner:2013vla, % Lin:2013sca, % Yu:2013aca, % Berlin:2014cfa % }, indirect detection \cite{ Beltran:2008xg, % Goodman:2010qn, % Cheung:2010ua, % Cheung:2011nt, % Rajaraman:2012fu, % DeSimone:2013gj, % Zheng:2010js, % Rajaraman:2012db, % Cheung:2012gi % } and direct detection \cite{ Belanger:2008sj,Kurylov:2003ra,Fitzpatrick:2012ib,Anand:2013yka,Fan:2010gt,Freytsis:2010ne,Cohen:2010gj,Gresham:2014vja,Fitzpatrick:2012ix } bounds on dark matter. The choice of pairwise dark matter interactions assumes the existence of a symmetry that also stabilizes the \DM particle against decay while the pairwise \SM interactions are assumed to be the leading order gauge-invariant operators. This need not be the case as has been demonstrated for annihilation \cite{Hochberg:2014dra} and direct detection \cite{Curtin:2013qsa}. In these cases, the structure in (\ref{eq:DM:EFT}) fails to capture the physics of the mediator fields which couple to both the dark and visible (\SM) sectors: the effective contact interaction description breaks down when the mediators do not decouple. The limitations of the contact interaction bounds were pointed out in \cite{Bai:2010hh} and highlighted in \cite{ Papucci:2014iwa, % Goodman:2011jq, % Busoni:2013lha, % Buchmueller:2013dya % }. This motivates a shift in the \emph{lingua franca} used to compare experimental results to models: rather than contact interactions, light (nondecoupled) mediators suggest using `simplified models' that include the renormalizable dynamics of the mediator fields \cite{Alves:2011wf}. This approach has been applied to colliders \cite{ Fox:2012ru, % Shoemaker:2011vi, % Friedland:2011za, % Graesser:2011vj, % An:2012va, % Frandsen:2012rk, % Profumo:2013hqa, % An:2013xka, Busoni:2013lha, % Buchmueller:2013dya, % Goodman:2011jq, % DiFranzo:2013vra, % Chang:2013oia, % Cotta:2013jna % } and astrophysical bounds where the physics of the mediator has been explored in \DM self-interactions \cite{Kaplinghat:2013kqa,Bellazzini:2013foa,Fan:2013yva,Tulin:2013teo,Tulin:2012wi,CyrRacine:2012fz,Foot:2012ai,Tulin:2012uq,Kaplan:2011yj,Loeb:2010gj,An:2009vq,Buckley:2009in,Feng:2009hw,Feng:2009mn,Pospelov:2008jd,ArkaniHamed:2008qn,Kesden:2006zb,Foot:2004pa,Mohapatra:2001sx, Spergel:1999mh}. \subsection{The $\gamma$-ray Excess Suggests Light Mediators} \label{sec:wby:light:mediators} \begin{table}[t] \begin{center} \begin{tabular}{ccl} \toprule % \textbf{Name} & \textbf{Operator} & \textbf{Constraint}\\ \rule{0pt}{3ex}% D2 & $(\bar\chi\gamma_5\chi) \, (\bar q q)$ & Edge of \EFT validity from monojet bounds \\ D4 & $(\bar\chi\gamma_5\chi) \, (\bar q \gamma_5 q)$ & Edge of \EFT validity from monojet bounds \\ \rule{0pt}{3ex}% D5 & $(\bar\chi\gamma^\mu\chi) \, (\bar q \gamma_\mu q)$ & Spin independent direct detection \\ D6 & $(\bar\chi\gamma^\mu\gamma_5\chi) \, (\bar q\gamma_\mu q)$ & Related to D5, D8 in chiral basis \\ D7 & $(\bar\chi\gamma^\mu\chi) \, (\bar q \gamma_\mu\gamma_5 q)$ & Related to D5, D8 in chiral basis \\ D8 & $(\bar\chi\gamma^\mu\gamma_5\chi) \, (\bar q \gamma_\mu\gamma_5 q)$ & Spin dependent direct detection \\ \rule{0pt}{3ex}% D9 & $(\bar\chi\sigma^{\mu\nu}\chi) \, (\bar q\sigma_{\mu\nu} q)$ & Nontrivial spin-2 \UV completion \\ D10 & $(\bar\chi\sigma^{\mu\nu}\gamma^5\chi) \, (\bar q\sigma_{\mu\nu} q)$ & Nontrivial spin-2 \UV completion \\ \rule{0pt}{3ex}% D12 & $(\bar\chi\gamma_5\chi) \, G_{\mu\nu}G^{\mu\nu}$ & Monojet bounds \\ D14 & $(\bar\chi\gamma_5\chi) \, G_{\mu\nu}\tilde G^{\mu\nu}$ & Monojet bounds \\ \bottomrule % \end{tabular} \end{center} \vspace{-1em} \caption{Contact operators between Dirac \DM and quarks or gluons \cite{Goodman:2010ku} that support $s$-wave annihilation and the constraint for the galactic center. See \cite{Alves:2014yha} for a recent technical analysis.} \label{tab:EFT:operators} \end{table} When the galactic center signal is combined with complementary bounds from direct detection and colliders, one is generically led to the limit where the contact interaction description (\ref{eq:DM:EFT}) breaks down and a simplified model description is necessary. By `generic' we mean no parameter tuning or additional model building is invoked. The tension is summarized in Table~\ref{tab:EFT:operators}, where we list the Dirac fermion dark matter contact interactions that satisfy the requirement of $s$-wave annihilation\footnote{% Majorana dark matter relaxes these bounds by forcing some of these operators to vanish identically. }. Because each effective operator simultaneously encodes the various \DM--\SM interactions in Fig.~\ref{fig:blob:diagrams}, requiring a coupling large enough to produce the $\gamma$-ray excess automatically generates signals that are constrained by null results at direct detection \cite{Akerib:2013tjd, Aprile:2013doa} and monojet \cite{ATLAS-CONF-2012-147} experiments. These rule out operators D5, D8, D12, and D14 in Table~\ref{tab:EFT:operators}. The operators D2 and D4 are at the edge of the validity of the effective theory \cite{ Busoni:2013lha, % Buchmueller:2013dya, % Goodman:2011jq}. % We ignore the D9 and D10 operators since they cannot be \UV completed by a renormalizable theory. Finally, the D6 and D7 operators are related to D5 and D8 by the chiral structure of the Standard Model. The fermionic SU(2)$_\text{L}\times$U(1)$_\text{Y}$ eigenstates are chiral so that gauge invariant interactions are naturally written in a chiral basis $\bar q \mathcal O_q P_{L,R} q$ where $P_{L,R}=\frac 12 (1\mp \gamma^5)$. Thus one generically expects that in the absence of tuning\footnote{It is worth noting that such a `coincidental' cancellation occurs in the $Z$ coupling to charged leptons which is dominantly axial due to $\sin^2\theta_W \approx 1/4$.\label{foot:lepton:Z:axial:coupling}}, the presence of vector or axial couplings implies the existence of the other. It is thus difficult to account for the $\gamma$-ray excess in the `heavy mediator' limit where these contact interactions are valid. A more technical analysis of the contact interaction description was recently performed in \cite{Huang:2013apa, Cheung:2011nt, Alves:2014yha} and includes the case of scalar dark matter. The $\gamma$-ray excess thus generically implies a dark sector with mediators that do not decouple and hence is more accurately described in a simplified model framework. Recent comprehensive studies of simplified models for the $\gamma$-ray excess have dark matter annihilating through off-shell mediators ($s$- and $t$-channel diagrams) \cite{Berlin:2014tja, Izaguirre:2014vva}; see \cite{Boehm:2014hva, Hektor:2014kga} for an earlier model. \subsection{Annihilation to On-shell Mediators} In this paper we focus on a different region in the space of simplified models where mediators are light enough that they can be produced on-shell in dark matter annihilation, henceforth referred to as the on-shell mediator scenario. This annihilation mode is largely independent of the mediator's coupling to the \SM so long the latter is nonzero. Lower limits on the \SM coupling---that is, upper limits on the mediator lifetimes---are negligible since the mediator may propagate astrophysical distances before decaying to the $b\bar b$ pairs that subsequently yield the $\gamma$-ray excess. The \SM coupling can be parametrically small which suppresses the off-shell $s$-channel annihilation mode as well as the direct detection and collider signals. This is shown in Fig.~\ref{fig:complimentary:mediator}. \begin{figure} \centering \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_chichitoOSmed} \caption{} \end{subfigure} \qquad\quad \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_chiSMtochiSMscalar} \caption{} \end{subfigure} \qquad\quad \begin{subfigure}[b]{0.25\textwidth} \includegraphics[width=\textwidth]{fig_SMSMtochichiscalar} \caption{} \end{subfigure} \caption{(a) Annihilation, (b) Direct Detection, (c) Collider. \DM complimentarity for on-shell mediators; compare to Fig.~\ref{fig:blob:diagrams}. (a) The annihilation rate is independent of the mediator coupling to the Standard Model. (b) Direct detection remains 2-to-2, here $N$ is a target nucleon. (c) Colliders can search for the presence of the mediator independently of its \DM coupling. } \label{fig:complimentary:mediator} \end{figure} Because on-shell annihilation into mediators requires at least two final states\footnote{% One may also consider semi-annihilation processes $\chi_1\chi_2 \to \chi_3 (\text{mediator})$~\cite{D'Eramo:2010ep}. See~\cite{Boehm:2014bia} for a prototype model for the galactic center $\gamma$-ray excess. }, the resulting annihilation produces at least four $b$ quarks, as shown in Fig.~\ref{fig:complimentary:mediator}a. This, in turn, requires a heavier dark matter mass in order to eject $\approx$ 40 \GeV $b$ quarks from each annihilation to fit the $\gamma$-ray excess. This avoids the conventional wisdom that this excess requires 10 -- 40 \GeV dark matter. In the limit on-shell annihilation dominates, the total excess $\gamma$-ray flux is fit by a single parameter, the mediator coupling to dark matter. Once fit, this parameter determines whether the \DM may be a thermal relic. We remark that the spectrum is slightly boosted by the on-shell mediator; we address this below and explore possibilities where the mediator mass can be used as a handle to change the spectral features. The on-shell mediator limit thus separates the physics of mediators \SM and \DM couplings. The former can be made parametrically small to hide \DM from direct detection and collider experiments, while the latter can be used to independently fit indirect detection signals such as the galactic center $\gamma$-ray excess. Observe that these simplified models modify the standard picture of complementary \DM searches for contact interactions shown schematically in Fig.~\ref{fig:complimentary:mediator}. Annihilation now occurs through multiple mediator particles and is independent of the mediator coupling to the \SM. Direct detection proceeds as usual through single mediator exchange between \DM and \SM. Collider bounds, on the other hand, need not depend on the \DM coupling at all and can focus on detecting the mediator rather than the dark matter missing energy. In this paper we explore the phenomenology of on-shell mediator simplified models for the galactic center. This paper is organized as follows. In the following two sections we present the on-shell simplified models that generate the $\gamma$-ray excess and determine the range of dark sector parameters. We then assess in Section~\ref{sec:experimental:bounds} the extent to which the on-shell mediators must be parametrically hidden from direct detection and colliders. In Section~\ref{sec:relic:abundance} we discuss the viability of this scenario for thermal relics. We comment on the lessons for \UV models of dark matter in Section~\ref{sec:UV:discussion}. Appendix~\ref{app:spectral:shape} briefly describes plausible variants for generating $\gamma$-ray spectra with more diverse \SM final states. | \begin{table} \begin{center} \begin{tabular}{cccccccclcl} \toprule % & \multicolumn{2}{l}{\textsc{Mass} [\textsc{g}e\textsc{v}] } & \, & \multicolumn{2}{l}{\textsc{Interaction}} & \quad & \multicolumn{2}{l}{\textsc{Coupling}} & & \textbf{Thermal} \\ \textbf{Mediator} & $m_\chi$ & $m_\text{mes.}$ & & \textbf{{\small DM}} & \textbf{{\small SM}} & & $\lambda_\text\DM$ & $\lambda_\text\SM$ & & \textbf{Relic?} \\ \hline \rule{0pt}{2ex} spin-0 & 110 & 20 & & $\gamma^5$ & $\mathbbm{1}$ & & 1.2 & $< 0.08$ % & & \textsc{msp}? \\ \textquotedbl & \textquotedbl & \textquotedbl & & $\gamma^5$ & $\gamma^5$ & & \textquotedbl & $<0.02^*$ & & \textquotedbl \\ spin-1 & 45 & 14 & & $\gamma^\mu$ & $\gamma_\mu$ & & 0.18 & % $< 10^{-6}$ & & $\gamma=1.3$ \\ \textquotedbl & \textquotedbl & \textquotedbl & & $\gamma^\mu\gamma^5$ & $\gamma_\mu\gamma^5$ & & \textquotedbl & $< 0.004$ &&\textquotedbl \\ \textquotedbl & \textquotedbl & \textquotedbl & & $\gamma^\mu\gamma^5$ & $\gamma_\mu$ & & \textquotedbl & $< 0.006$ &&\textquotedbl \\ \textquotedbl & \textquotedbl & \textquotedbl & & $\gamma^\mu$ & $\gamma_\mu\gamma^5$ & & \textquotedbl & $< 0.02$ &&\textquotedbl \\ \bottomrule % \end{tabular} \caption{Best fit parameters assuming $b$-philic couplings for the spin-0 mediator and universal quark couplings for the spin-1 mediator. The upper bound for $\lambda_\text\SM$ for the $\gamma^5\otimes\gamma^5$ is a conservative estimate for the 8 \TeV mono-$b$ reach at the \LHC (see Section~\ref{sec:collider}); the other bounds come from direct detection. In the last column, we indicate whether consistency with a thermal relic abundance suggests a tighter \DM profile ($\gamma=1.3$) or some population of millisecond pulsars (\textsc{msp}), see Section~\ref{sec:relic:abundance}. }\label{table:conclusion} \end{center} \end{table} We have presented a class of simplified models where dark matter annihilates into on-shell mediators which, in turn, decay into the \SM with a typically suppressed width. This separates the sector of the model which can account for indirect detection signals---such as the \FERMI galactic center $\gamma$-ray excess---and those which are bounded by direct detection and collider experiments. We have addressed $\gamma$-ray spectrum coming from these models and have compared used the $\gamma$-ray excess to identify plausible regions of parameter space for a \DM interpretation; the best fit parameters and bounds on the \SM coupling are shown in Table~\ref{table:conclusion}. We have addressed the key points for \UV model building and, in an appendix below, highlight further directions for modifying the $\gamma$-ray spectrum with more general \SM final states. | 14 | 4 | 1404.6528 |
1404 | 1404.2417_arXiv.txt | The anisotropic nature of active galactic nuclei (AGN) is thought to be responsible for the observational differences between type-1 (pole-on) and type-2 (edge-on) nearby Seyfert-like galaxies. In this picture, the detection of emission and/or absorption features is directly correlated to the inclination of the system. The AGN structure can be further probed by using the geometry-sensitive technique of polarimetry, yet the pairing between observed polarization and Seyfert type remains poorly examined. Based on archival data, I report here the first compilation of 53 estimated AGN inclinations matched with ultraviolet/optical continuum polarization measurements. Corrections, based on the polarization of broad emission lines, are applied to the sample of Seyfert-2 AGN to remove dilution by starburst light and derive information about the scattered continuum alone. The resulting compendium agrees with past empirical results, i.e. type-1 AGN show low polarization degrees ($P~\le$~1~\%) predominantly associated with a polarization position angle parallel to the projected radio axis of the system, while type-2 objects show stronger polarization percentages ($P~>$~7~\%) with perpendicular polarization angles. The transition between type-1 and type-2 inclination occurs between 45$^\circ$ and 60$^\circ$ without noticeable impact on $P$. The compendium is further used as a test to investigate the relevance of four AGN models. While an AGN model with fragmented regions matches observations better than uniform models, a structure with a failed dusty wind along the equator and disc-born, ionized, polar outflows is by far closer to observations. However, although the models correctly reproduce the observed dichotomy between parallel and perpendicular polarization, as well as correct polarization percentages at type-2 inclinations, further work is needed to account for some highly polarized type-1 AGN | \label{Intro} The unified model of active galactic nuclei (AGN; \citealt{Antonucci1993}) states that most of the observational differences between type-1 and type-2 Seyfert-like galaxies arise from an orientation effect. According to this theory, the disappearance of ultraviolet (UV) and optical broad emission features at type-2 inclinations can be explained by the presence of an obscuring, circumnuclear material along the equatorial plane of the AGN (the so-called dusty torus) that hides both the central engine and the photoionized broad line regions (BLRs; the low ionization line LIL BLR and the highly ionized line HIL BLR; \citealt{Rowan1977,Osterbrock1978}). A type-2 viewing angle can then be defined as a line of sight towards the central source that intercepts the equatorial dust, while type-1 inclination allows a direct view of the central engine. The observational lack of type-1 AGN with edge-on host galaxy \citep{Keel1980,Lawrence1982} suggests that dust along Seyfert~1 galaxy discs may obscure the HIL and LIL BLR and make the AGN appear like Seyfert~2s \citep{Maiolino1995}. The number count of Seyfert~1 objects is thus expected to be small, even if the fraction of type-1 against type-2 AGN in the nearby Universe still needs to be properly determined. Estimating the orientation of a large sample of AGN is necessary to verify the assumptions of the unified model, and check whether all the differences between Seyfert~1 and Seyfert~2 objects can be explained by inclination or if morphological differences must also be taken into account. In this regard, polarization has proven to be a solid tool to investigate the inner structure of AGN. The spectropolarimetric measurements of NGC~1068 by \citet{Miller1983} helped to identify electron and dust scattering as the main mechanisms producing a continuum polarization in radio-quiet AGN. Going further, the extensive high-resolution, high signal-to-noise spectropolarimetric observation of the same AGN by \citet{Antonucci1985} revealed the presence of highly polarized, broad, symmetric Balmer and permitted Fe~{\sc ii} lines. The polarization spectrum was found to be closely similar to typical Seyfert~1 galaxies, supporting the idea that Seyfert~2 AGN are hiding Seyfert~1 core behind the dusty torus. This discovery was a key argument in favour of a unified model of AGN. Spectropolarimetry is thus a powerful method to probe the validity of any AGN model, as the computed fluxes shall match both observational intensity, polarization percentage and polarization angle, reducing the number of free parameters/degeneracies \citep{Kartje1995,Young2000,Goosmann2007,Marin2012a}. In order to model a peculiar source, the observer's viewing angle $i$ has to be set (e.g. $i \sim$ 70$^\circ$ for NGC~1068; \citealt{Honig2007,Raban2009}) to explore the resulting polarization \citep{Goosmann2011,Marin2012c}. The impact of the system's orientation on to the net polarization can lead to significantly different results, especially when the observer's line of sight matches the half-opening angle of the obscuring region \citep{Marin2012a}. To be consistent with observation, an investigation of the model over a broad range of inclination must be undertaken. However, any comparison between the observed polarization and the theoretical orientation of individual AGN is hampered by the lack of a data base that combines inclination and polarization. It is the aim of this paper to provide the first spectropolarimetric compendium of Seyfert-like galaxies, gathering observed continuum polarizations from literature correlated with estimated inclinations of individual AGN. In Sect.~\ref{Comp}, I investigate the different observational and numerical techniques used to estimate the inclination of 53 objects, and match the sources with their archival UV/optical polarization measurement, whenever it is feasible. To illustrate the significance of this catalogue when comparing models to observations, in Sect.~\ref{Analysis}, I pick four different, competitive AGN models from the literature and analyse them in the framework of this compendium. In Sect.~\ref{Discussion}, I review the successes and potential improvements of AGN modelling, explore the problematic, high polarization levels of peculiar type-1 objects and discuss possible bias on the estimation of inclination. Finally, conclusions are drawn in Sect.~\ref{Conclusion}. | \label{Conclusion} The first match of 53 AGN inclinations with their intrinsic continuum polarization originating from electron and dust scattering is presented in this paper. Different techniques to retrieve the nuclear orientation of type-1 and type-2 Seyfert galaxies were presented and discussed, highlighting their potential caveats. The continuum polarization of several Seyfert-2s was corrected using broad H$\alpha$ and H$\beta$ line polarization as a reliable indicator of the true polarization of the scattered light, and lower limits were put for the remaining AGN whose polarization spectra were either noise-saturated or unpublished. ~\ The resulting compendium\footnote{The compendium will be regularly updated and available upon email request.} is in agreement with past observational/theoretical literature, and warrants additional conclusions and remarks the following. \begin{itemize} \item Seyfert 1 AGN are associated with low polarization degrees, $P~\le$~1~\%, and predominantly characterized by a polarization position angle parallel to the projected radio axis of the system. The inclination of type-1 objects ranges from 0$^\circ$ to 60$^\circ$. \item Seven type-1s have been identified as polar scattering dominated AGN, i.e. showing a perpendicular polarization position angle. Among them, five have an atypical continuum polarization higher than 1~\%, mostly associated with 10$^\circ$ -- 45$^\circ$ inclinations. As scattering-induced polarization is unlikely to produce such high polarization degrees at type-1 orientation, a more elaborate scenario must be considered. \item After correction, Seyfert 2 AGN show polarization degrees higher than 7~\% and perpendicular polarization position angle. Unfortunately, most of the objects have only lower limits. The inclination of type-2 objects is ranging from 47$^\circ$ to 90$^\circ$. \item The transition between type-1 and type-2 AGN occurs between 45$^\circ$ -- 60$^\circ$. This range of inclination is likely to include AGN classified as borderline objects, where the observer's line of sight crosses the horizon of the equatorial dusty medium. Four objects lie in this range, three of them (ESO~323-G077, NGC~1365 and IRAS~13349+2438) being already considered as borderline Seyfert galaxies from spectroscopic observations. If the estimated inclination of the fourth object (the Circinus galaxy) is correct, it should be considered as another borderline AGN. \item The usual axisymmetric AGN models have difficulties to reproduce the trend of polarization with inclination. Fragmenting the reprocessing regions is helpful to cover a wide range of continuum polarization but a disc-born wind model is found to be already quite close from observations.A fine tuning of the line-driven disc wind could easily match a substantial fraction of the reported measurements. \end{itemize} ~\ Problems determining the inclination of AGN must be taken into consideration but, despite potential caveats, the associated continuum polarization lies within the margins of past empirical results, consolidating the basis of the compendium. It is then important for future models, as a consistency check, to reproduce the average continuum polarization, the polarization dichotomy and the transition between type-1 and type-2 classification (45$^\circ < i <$ 60$^\circ$). By improving the quality of the methods to determine the inclination of AGN and properly removing the contribution of both stellar and starburst light in future polarimetric measurement of type-2 objects, it will possible to bring very strong constraints on the morphology, composition and kinematics of AGN. To achieve this goal, a new UV/optical spectropolarimetric atlas of Seyfert 2s is necessary. | 14 | 4 | 1404.2417 |
1404 | 1404.2883_arXiv.txt | Two component advective flows are the most physical accretion disks which arise from theoretical consideration. Since viscosity is the determining factor, we investigate the effects of viscous stresses on accretion flows around a nonrotating black hole. As a consequence of angular momentum transfer by viscosity in an accretion flow, the angular momentum distribution is modified. We include cooling effects and found that a Keplerian disk is produced on the equatorial plane and the flow above and below remains sub Keplerian. This gives a complete picture to date, of the formation of a Two component advective flow around a black hole. | Observation evidence of non-thermal photons in the spectrum (Sunyaev \& Truemper, 1979) prompted the model developers to think that a hot electron cloud (the so-called Compton cloud) along with the standard disk could resolve the issue (Sunyaev \& Titarchuk, 1980). Numerous suggestions and cartoon diagrams of the illusive The presence of Compton cloud are shown in the literature (e.g., Zdziarski, 1988 ; Haardt et al., 1994 ; Chakrabarti \& Titurchuk, 1995 ; Hereafter CT95). CT95, based on the solutions of viscous and inviscid transonic flows around black holes (Chakrabarti, 1989; Chakrabarti 1990) proposed that, in general, the accretion disk should really have two components: a Keplerian accretion on the equatorial plane and a sub-Keplerian halo which surrounds the Keplerian disk, and the puffed up inner part of the flow (CENBOL) which is nothing but the Compton cloud. There was as yet no work in the literature to show that the Two Component Advective Flow (TCAF) solution is stable. The cause for concern was obvious: a Keplerian disk is necessarily sub-sonic, while the sub-Keplerian flow is supersonic, and becomes sub-sonic only at the shock wave. Thus the questions remained unanswered is: Under what circumstances TCAF actually forms? In this paper, through numerical simulations of viscous accretion flow with a power-law cooling effects, we show that when the injected sub-Keplerian flow angular momentum is high enough and/or the viscosity is high enough, TCAF would be formed, otherwise the sub-Keplerian flow would remain sub-Keplerian. | So far, there was no numerical simulations in the literature to show that TCAF solution is realizable as a whole, and there was no simulation to show whether such a configuration is at all stable. Our result, for the first time, shows that if one assumes that the viscosity is maximum on the equatorial plane, then, a low- angular momentum injected flow is converted into a TCAF. We show that the injected flow is segregated in the Keplerian and the sub-Keplerian components. So far, we have captured all the salient features of the Keplerian disk, by introducing a power-law cooling effect. In order to produce an exact standard disk which emits multicolour black body as well, we need to include the radiative transfer problem. | 14 | 4 | 1404.2883 |
1404 | 1404.2284_arXiv.txt | Observations indicate that our universe is characterized by a late-time accelerating phase, possibly driven by a cosmological constant $\Lambda$, with the dimensionless parameter $\Lambda L_P^2 \simeq 10^{-122}$, where $L_P = (G \hbar /c^3)^{1/2}$ is the Planck length. In this review, we describe how the emergent gravity paradigm provides a new insight and a possible solution to the \cc\ problem. After reviewing the necessary background material, we identify the necessary and sufficient conditions for solving the \cc\ problem. We show that these conditions are naturally satisfied in the emergent gravity paradigm in which (i) the field equations of gravity are invariant under the addition of a constant to the matter Lagrangian and (ii) the cosmological constant appears as an integration constant in the solution. The numerical value of this integration constant can be related to another dimensionless number (called \cn) that counts the number of modes inside a Hubble volume that cross the Hubble radius during the radiation and the matter dominated epochs of the universe. The emergent gravity paradigm suggests that \cn\ has the numerical value $ 4 \pi$, which, in turn, leads to the correct, observed value of the cosmological constant. Further, the emergent gravity paradigm provides an alternative perspective on cosmology and interprets the expansion of the universe itself as a quest towards holographic equipartition. We discuss the implications of this novel and alternate description of cosmology. | Several recent results suggest that the field equations of gravity have the same conceptual status as the equations of, say, elasticity or fluid mechanics, making gravity an emergent phenomenon (for a review, see \cite{tpreviews}). This paradigm provides an alternative description of spacetime evolution as being driven by the difference between the surface and bulk degrees of freedom, and can be derived from a thermodynamic variational principle that extremizes the heat content of null surfaces \cite{grtd}. This approach also provides a strikingly different perspective on cosmology in general and the \cc\ in particular. The main goal of this review is to explain how one can obtain a solution to the cosmological constant problem in the backdrop of the emergent perspective of gravity. The plan of the review is as follows: In the next section, we start with a brief overview of the conventional approach to cosmology in which the Friedmann universe is described in terms of the standard parameters like $H_0,\om,\orr,...$ etc. We then motivate and describe an alternative, \textit{epoch invariant} parameterization of cosmology in which the universe is described by a set of variables (essentially three energy densities, $\rho_{\rm inf},\rho_{\rm eq}$ and $\rho_{\Lambda}$) whose numerical values are manifestly independent of the epoch at which they are measured. We identify these as the three energy densities associated with the three phases of evolution of the universe, viz. the inflationary phase ($\rho_{\rm inf}$), the radiation-matter dominated phase ($\rho_{\rm eq}$) and the late time accelerating phase ($\rho_{\Lambda}$). Of these three, $\rho_{\rm inf}$ and $\rho_{\rm eq}$ are expected to be determined by high-energy physics models. However, there does not exist a physical principle which determines the extremely small value of $\rho_\Lambda L_P^4 \approx 10^{-123}$ in natural units. After briefly reviewing (Sec.\ref{sec:approaches}) several approaches in the literature to handle the above cosmological constant problem, we identify, in Sec.\ref{sec:nscc}, the essential theoretical ingredients which are needed for this purpose. We argue that, for a viable solution to the problem, we must have the following three features in the theory of gravity: (a) The field equations of gravity must be invariant under the addition of a constant to the matter Lagrangian. (b) The cosmological constant must appear as an integration constant in the solutions. (c) To determine its value, a new physical principle is required. Of these three requirements, (a) and (b) are known to be naturally satisfied in the emergent gravity paradigm, as described in Sec.\ref{sec:egandcc}. We, therefore, start with the requirement (c) in Sec.\ref{sec:cosmin}. Using the epoch invariant parametrization of cosmology, we construct a dimensionless number (which we call \cn), that counts the number of modes within a Hubble volume that cross the Hubble radius from the end of inflation to the beginning of the late-time accelerating phase. We show how the introduction of \cn\ and the postulate that the numerical value of \cn\ is $4 \pi$ allow us to determine the \textit{correct, observed value} of the cosmological constant for a GUTs scale inflation and the allowed range in the matter and radiation energy densities as determined from cosmological observations. In other words, \cn\ allows the determination of the numerical value of the cosmological constant from first principles in terms of other parameters that are expected to be determined from high energy physics. In Sec.\ref{sec:egandcc}, we examine these issues in the backdrop of the emergent gravity paradigm, in which the evolution of the universe is described as a quest towards holographic equipartition, and the field equations of gravity are invariant under the additional of a constant to the matter Lagrangian with the cosmological constant appearing as an integration constant in the solutions. Together with the introduction of \cn\ and the postulate that \cn\ $= 4 \pi$, this provides a comprehensive approach that determines the value of the cosmological constant. We use the signature $(-, +,+,+)$ and units with $c = \hbar = 1$ unless otherwise specified. All numerical values of cosmological parameters are consistent with the Planck 2013 data \cite{planck2013}. | The dynamics of the spatially flat universe can be characterized by three densities: $\rho_{\rm inf},\rho_{\rm eq}$ and $\rho_\Lambda$, in addition to one undetermined overall normalization constant for $a(t)$ which could be taken to be the value of the expansion factor at the epoch when $\rho_m=\rho_R$. In standard cosmology, there is no inter-relationship between the three densities $\rho_{\rm inf},\rho_{\rm eq}$ and $\rho_\Lambda$. It is generally believed that high energy physics models will eventually provide a first-principles estimation of both $\rho_{\rm inf}$ and $\rho_{\rm eq}$. But we have no clue so far as to which physical principle determines the value of $\rho_{\Lambda}$. We have shown that there exists a characteristic number $N_c$ for the universe (``\cn'') which counts the number of length scales which enter the Hubble radius during the radiation/matter dominated phase. We have argued in Sec.\ref{sec:cosmin} that this conserved number can act as a unifying link between the early inflationary phase, the radiation/matter dominated phase and the late-time acceleration phase. For a generic universe described by an unrelated set of three densities $\rho_{\rm inf},\rho_{\rm eq}$ and $\rho_\Lambda$, the parameter \cn\ can take any arbitrary value. \textit{One of our key results is the discovery that \cn\ for our universe is equal to $4\pi$ to a high degree of accuracy.} (See \eq{muvalue}.) We strongly believe that this is unlikely to be an accident and demands an explanation. If this result, $N_c=4\pi$ is raised to a status of a postulate, it can be used to determine the numerical value of $\rho_\Lambda$ in terms of the other two densities $\rho_{\rm eq}$ and $\rho_{\rm inf}$ (see \eq{ll6}) which --- as we said before --- are very likely to be determined by high energy physics. In other words, we now have a paradigm in hand in which all the numbers characterizing the universe are determined from first principles. There is no a priori reason for such an idea to give results consistent with observations, which, in fact, happens only for a narrow range of inflationary energy scales. As mentioned in the first item in Sec.\ref{sec:disc}, this range is likely to be constrained only by a factor of about 15 even if we take into account uncertainties due to inflationary reheating. In any case, given a detailed model of inflation, the procedure for calculating \cn\ is well defined and one could check for the consistency of the idea when inflationary models are on firmer footing. Such a procedure to determine the numerical value of $\rho_\Lambda$ has deeper implication for the structure of gravitational theories. In conventional models, the value of $\rho_\Lambda$ can change when an arbitrary constant is added to the matter Lagrangian. In such a case, any physical principle to determine the numerical value of $\rho_\Lambda$ becomes dubious, since the value of the \cc\ that acts as the source for gravity can be changed by adding a constant to the matter Lagrangian. As we have pointed out in Sec.\ref{sec:nscc}, this \textit{is a generic problem} in a large class of attempts to ``solve'' the \cc\ problem and the \cc\ problem can be really solved \textit{only if} we have two separate key ingredients in our model: (a) The gravitational \textit{field equations} must be invariant under the addition of a constant to the matter Lagrangian (which results in the modification of the zero level of energy, as in \eq{tsym}). (b) At the same time, the \textit{solutions} to the gravitational field equations must allow for the inclusion of a \cc, and we must provide a new principle to determine its numerical value. It was known for a long time \cite{aseemtp,tpap} that the emergent gravity paradigm, leading to the field equations of the form in \eq{tf}, takes care of the ingredient (a) above. What was lacking was a new physical principle which could be used to determine the value of the \cc, which arises as an integration constant to the solution of \eq{tf}. Here, we have provided this second ingredient (b) in the form of our postulate $N_c = 4\pi$. As mentioned in Sec.\ref{sec:coscon}, it is difficult to reconcile a result like $N_c=4\pi$ within the conventional cosmological paradigm which treats the dynamics of the universe simply as a solution to the gravitational field equations. In fact, the conventional approach does not provide satisfactory answers to several other conceptual questions one could raise about our universe. For example, one cannot even answer a simple question as to ``why does the universe expand'' within the context of standard classical cosmology \cite{TPwhy}. Since gravitational field equations are invariant under time reversal, one can write down the time-reversed solutions to the Friedmann equations describing a contracting universe. All that we can prove is that \textit{if} the universe is expanding today (which is taken as an observational input) \textit{then} it would have been expanding in the past --- though the initial singularity prevents us from meaningfully setting ``initial'' conditions to choose this solution. Further, in conventional cosmology, the universe seems to have evolved \textit{spontaneously} from a quantum mechanical state to a nearly classical state. It is not possible for a normal system to make a \textit{spontaneous} transition from a quantum to a classical state (in a rigorously defined sense, in terms of the Wigner function) if its evolution is governed by a bounded Hamiltonian \cite{TPwhy}. Finally, the Friedmann model breaks Lorentz invariance and chooses a preferred Lorentz frame (in which the CMB is isotropic), again because the solution to the field equations breaks the full symmetry of the field equations. All these features seem to suggest \cite{tpbj2} that we should describe cosmic expansion (and derive the Friedmann equations) from another physical principle, rather than treat them as arising as a solution to the gravitational field equations. We have argued in Sec.\ref{sec:coscon} that the concept of holographic equipartition provides such an alternate principle to describe cosmic expansion. This concept uses the Hubble radius fairly crucially and also identifies $4\pi$ as a primordial constant counting the number of degrees of freedom on a sphere of radius $L_P$. Both these ingredients go well with $N_c= 4\pi$ acting as a fundamental physical principle. Though the ideas presented in Sec.\ref{sec:coscon} of the review are still at a preliminary stage compared to the rest of the review, but they hold the promise of providing a novel and fruitful description of our cosmos. | 14 | 4 | 1404.2284 |
1404 | 1404.0248_arXiv.txt | In this paper we analyze the peculiar radio structure observed across the central region of the galaxy cluster Abell\,585 ($z=0.12$). In the low-resolution radio maps, this structure appears uniform and diffuse on angular scales of $\sim 3$\arcmin, and is seemingly related to the distant ($z=2.5$) radio quasar B3\,0727+409 rather than to the cluster itself. However, after a careful investigation of the unpublished archival radio data with better angular resolution, we resolve the structure into two distinct arcmin-scale features, which resemble typical lobes of cluster radio galaxies with no obvious connection to the background quasar. We support this conclusion by examining the spectral and polarization properties of the features, demonstrating in addition that the analyzed structure can hardly be associated with any sort of a radio mini-halo or relics of the cluster. Yet at the same time we are not able to identify host galaxies of the radio lobes in the available optical and infrared surveys. We consider some speculative explanations for our findings, including gravitational wave recoil kicks of SMBHs responsible for the lobes' formation in the process of merging massive ellipticals within the central parts of a rich cluster environment, but we do not reach any robust conclusions regarding the origin of the detected radio features. | During merging processes leading to the formation of clusters of galaxies, large amounts of gravitational energy are released on timescales of the order of $\sim$\,Gyr. Most of this energy is contained in hot (temperatures $kT \lesssim 10$\,keV) X-ray--emitting plasma which constitutes, along with the dark matter, the dominant fraction of the intracluster medium (ICM; e.g., \citealt{sa1}). In addition to the thermal gas, however, $\sim \mu$G magnetic fields and ultrarelativistic electrons are present within the ICM as well, manifesting most clearly in extended diffuse radio structures such as giant and mini radio halos in the central parts of clusters, or radio relics at cluster peripheries (see the reviews by \citealt{ct02,ferrari08,feretti1}). These non-thermal constituents of the ICM are believe to be related to the energy dissipation processes enabled by large-scale shocks formed at the outskirts of merging systems, and/or turbulence driven by various mechanisms at early post-merger stages of a cluster lifetime. The presence of (or rather the amount of) hadronic cosmic rays in the diffuse cluster environment is still an open question, being currently probed with new-generation $\gamma$-ray instruments (see, e.g., \citealt{ah09,ac10,al10}). Besides radio halos and relics, extended lobes and plumes of radio galaxies make up yet another class of diffuse non-thermal structures often found in the ICM. These are formed due to the jet activity of accreting supermassive black holes (SMBHs) hosted by the cluster galaxies. The large-scale morphologies of such lobes are rarely regular or symmetric, unlike in the case of `classical doubles' typical for poorer (galaxy group) environments. Instead, lobes of cluster radio galaxies are often bent (`tailed' radio sources), irregular, or even amorphous, reflecting the dramatic impact of high-pressure ambient medium, high peculiar velocities of parent galaxies, and highly intermittent jet activity of cluster SMBHs (e.g., \citealt{mil80}; see also the discussion in \S\,5.3 below). Such structures may therefore be considered as useful probes of the dynamical state and structure of the ICM, or even as effective tracers of galaxy clusters at moderate and high redshifts (see, e.g., \citealt{mao1}). Since they are best characterized at low radio frequencies, the operation of the next generation of low-frequency interferometers like the LOw Frequency ARray \citep{vH13} and the Murchison Widefield Array \citep{tin13} are in this context much anticipated. Recently, much attention has been given to radio sources located within the central parts of galaxy clusters, since the mechanical energy output of radio-loud active galactic nuclei (AGN) which are hosted by central elliptical (cD) galaxies is widely believed to be responsible for suppressing the cluster cooling flows, quenching the star-formation in cD systems and the growth of their SMBHs (see, e.g., \citealt{fa94,be04,mcn07}). In order to investigate the details of such feedback processes, a precise spectral and morphological characterization of central radio structures and of their surroundings is however obligatory, and this requires deep high-resolution exposures with various instruments operating from radio to X-ray bands. Without such extensive dataset, the physics of the lobes-ICM interactions may remain elusive. In this paper we analyze the peculiar radio structure observed in the direction of the central parts of the galaxy cluster Abell\,585 ($z=0.12$). The paper is organized as follows. The galaxy cluster Abell\,585 is briefly introduced in \S\,2. Multi-frequency data for the system are analyzed in \S\,3, and discussed further in \S\,4. The interpretation of the obtained observational results is presented in \S\,5, and the summary of the studies is given in \S\,6. In the paper we assume modern cosmology with $H_0=71$\,km\,s$^{-1}$\,Mpc$^{-1}$, $\Omega_{\rm M}=0.27$, and $\Omega_{\Lambda}=0.73$ \citep{s1}. All source positions are given in the J2000.0 coordinate system. | The peculiar arcmin-scale radio structure observed in the direction of the central parts of the galaxy cluster Abell\,585 ($z=0.12$) is unassociated with the distant ($z=2.5$) radio quasar B3\,0727+409, as we demonstrated here by means of a careful analysis of all the available radio data for the system. This structure consists of two dominant features which resemble typical lobes of cluster radio galaxies of the NAT type. However, we are not able to identify host galaxies of the two features in the available optical (SDSS) and infrared (WISE) surveys. We speculate if the analyzed systems are examples of extreme gravitational recoils of SMBHs in the process of merging massive ellipticals within the central parts of a rich cluster environment, but we do not reach any robust conclusions regarding the origin of the detected radio features. Large-scale radio structures lacking obvious host galaxies may not be that unique for the local clusters. For example, radio maps of B\,1753$+$580 found at the position of Abell\,2289 \citep{o0} show two edge-dimmed tails with bright `heads' and with a $z=0.224$ galaxy located between them \citep{ow1}, but without any detectable compact radio core. \citet{Slee98} presented a detailed study of the radio structure in cluster Abell\,4038 resembling the evolved lobes of a luminous radio galaxy, but with the most likely host elliptical galaxy displaced by about 20\,kpc with respect to the central parts of the radio structure. \citet{Slee01} discussed three other possibly analogous systems in clusters Abell\,13, Abell\,85, and Abell\,133. Inspecting FIRST maps of several galaxy clusters from \citet{koe07}, we add the case of Max BCG J130.34220+61.21246 ($z=0.13$) to the literature examples, and note the striking similarity between the radio structures discussed in this paper and those in the $z = 0.16$ cluster MaxBCG J250.34552+38.03597 that lack any optical identifications. There may be more examples of such peculiar structures awaiting discovery with high-resolution multiwavelength radio observations of lesser-known clusters of galaxies in formation. | 14 | 4 | 1404.0248 |
1404 | 1404.7156_arXiv.txt | The rotation rate, level of magnetic activity and surface lithium abundance are age-dependent quantities in stars of about a solar mass and below. The physical reasons for the evolution of these phenomena are qualitatively understood, but accurate quantitative models remain dependent on empirical calibration using the Sun and stars of known age, chiefly in clusters. In this work I review the status of these ``empirical age indicators'', outlining the astrophysics of their time dependence, describing the measurements, assessing the precision (and accuracy) of age estimates when applied to individual stars, and identifying their principle limitations in terms of the mass and age ranges over which they are useful. Finally, I discuss the ``lithium depletion boundary'' technique which, in contrast to the empirical methods, appears to provide robust, almost model-independent ages that are both precise and accurate, but which is only applicable to coeval groups of stars. | \label{sec1} The age of a star is, along with its mass and composition, the most important quantity to know for testing ideas concerning the evolution of stars, stellar systems (clusters and galaxies) and also, by association, their circumstellar material and exoplanetary systems. However, unlike mass and composition, we have no direct means of measuring the age of any star but the Sun. The ages of other stars are inferred or estimated using a hierarchy of techniques, which can be described as (see Soderblom 2010; Soderblom \etal\ 2013) semi-fundamental, model-dependent, empirical or statistical. Semi-fundamental techniques rely on age-dependent phenomena where the physics is understood, there is little tuning of model parameters required and the results are basically model-independent. Model-dependent techniques include isochrone fitting in the Hertzsprung-Russell (HR) diagram, asteroseismology and white dwarf cooling. Here the physics is mostly understood, but there are annoying gaps in our ability to accurately model the physics without making simplifying assumptions or tuning parameters (e.g. the mixing length) to match observations. Often the precision of the ages determined by such techniques is much better than their absolute accuracy and different models may yield ages that differ by more than their claimed uncertainties. At a level below the model-dependent techniques are empirical age indicators. For these, the understanding of the physics is qualitative, with significant holes in the theory that are usually bridged using semi-empirical relationships with free parameters. The general approach is to calibrate an age-dependent phenomena using similar observations of stars with ``known'' age (the Sun and stars with ages estimated by semi-fundamental or model-dependent techniques) and then use that calibration to estimate the ages of other stars (e.g. Barnes 2007; Vican 2012). Of course, there is a risk of circularity here; one cannot study the age dependence of a phenomenon using stars with ages estimated using that phenomenon! In this contribution I deal mainly with empirical age indicators associated with the rotation rates, levels of magnetic activity and photospheric lithium abundances of stars with masses $M\leq 1.3$\msun\ and how they apply to stars from birth to ages of 10\,Gyr. It is no coincidence that these phenomena each become useful below this mass. The presence of a sub-photospheric convection zone is responsible for dynamo-generated magnetic fields that are dissipated to provide non-radiative heating in the outer atmosphere and also couple to an ionised wind that drives angular momentum loss. The same convection zone is responsible for mixing pristine material down to interior regions where Li can be burned. The use of these indicators has its root in work done by Kraft and collaborators in the 1960s (e.g. Kraft \& Wilson 1965; Kraft 1967), but perhaps the most influential early paper was by Skumanich (1972), who showed that both rotation and activity, and to some extent Li abundance, decayed according to the inverse square root of age. The data used were sparse, consisting of the Sun (age 4.57\,Gyr) and a few solar-type stars in the Pleiades (age $\simeq 125$\,Myr) and Hyades (age $\simeq 600$\,Myr) open clusters, but nevertheless this paper stimulated much of what follows. The utility of these empirical age indicators is mostly in estimating ages for low-mass main sequence (MS) and pre main sequence (PMS) stars that constitute the vast majority of the Galactic population. A principle advantage of the techniques I will discuss is that they are {\it distance independent}. With the successful launch of the {\it Gaia} satellite (Perryman \etal\ 2001; Brown 2008), it might seem that uncertain stellar distance will be a solved problem within a few years. However, even with precisely known distances, the determination of ages for stars that have reached the main sequence and are still burning hydrogen in their cores is difficult. Position in the HR diagram is age sensitive, but also sensitive to the detailed composition of the star. Even with [Fe/H] known to a very respectable accuracy of $\pm 0.05$\,dex, the age of a 5\,Gyr solar-mass star could only be estimated to a precision of 20 per cent, and considerably worse for lower mass stars with longer main sequence lifetimes that consequently move more slowly in the HR diagram (e.g. see Fig.~20 of Epstein \& Pinsonneault 2014). Asteroseismology may be an alternative distance-independent method for age estimation, with the advantage of a strong and well-understood physical basis, but it is not clear that pulsations can easily be detected in main-sequence stars well below a solar mass or in young, active stars (e.g. Huber \etal\ 2011). Even if they are, it is unlikely that ages could presently be estimated for solar-type stars to absolute precisions better than 10--15 per cent of their main sequence lifetimes (e.g. Gai \etal\ 2011; Chaplin \etal\ 2014) and would rapidly become too large to be useful in stars below a solar mass. Hence, there is likely to be a need for age determinations using empirical indicators for the forseeable future. In section~\ref{sec2} I discuss measurements of rotation in low-mass stars, the physical basis on which rotation rate could be used to estimate age and review efforts to calibrate ``gyrochronology''. Section~\ref{sec3} reviews the connection between rotation and magnetic activity and the various attempts to calibrate activity-age relationships using several magnetic activity indicators. Section~\ref{sec4} discusses the astrophysics of lithium depletion in solar-type stars, comparison of observations and models and the use of lithium as an empirical age indicator in PMS and MS stars separately. Also included is a description of the ``lithium depletion boundary'' technique in very low mass stars, which differs from the other methods discussed here in that it requires no empirical calibration and is semi-fundamental. Section~\ref{sec5} summarises the status and range of applicability of each of these techniques and briefly discusses efforts to improve empirical calibrations. Conclusions are presented in section~\ref{sec6}. | The need for empirical methods of age estimation in low-mass stars is likely to be present for some years to come. In this contribution I have reviewed the astrophysical reasons that rotation, magnetic activity and the photospheric abundance of lithium, change with time in low-mass stars ($\leq 1.3\,M_{\odot}$). Whilst theoretical models that predict these behaviours are improving rapidly, there are still very significant uncertainties and semi-empirical components that prevent their use in directly estimating stellar ages with any certainty, and which require calibration using stars of known age. Each of these empirical age indicators can play a role in various domains of mass and age, that are schematically illustrated in Fig.~\ref{summaryplot}. The rotation-age relationship (or gyrochronology) offers the best prospect of determining precise (to 20 per cent) ages in older stars, and could be complemented by the use of PMS Li depletion to estimate the ages of younger stars at low masses. Magnetic activity offers a less precise age determination in older stars, but is usually easier to measure than rotation. In terms of accuracy, all these methods are compromised to some extent by a lack of calibrating data in stars that are older than the Sun or of lower masses than the Sun. In very low mass stars, the sharp transition between stars that have depleted all their lithium and stars with similar age but only slightly lower luminosities that have preserved all their lithium (the lithium depletion boundary), offers an almost model-independent way of estimating an age for groups of coeval stars. This technique is sensitive between ages of 20 and 200\,Myr and can be used to investigate the uncertain physics in stellar models or calibrate empirical age indicators. \label{sec6} | 14 | 4 | 1404.7156 |
1404 | 1404.1400_arXiv.txt | We study the viability of pseudo Nambu-Goldstone bosons (Majorons) arising in see-saw models as dark matter candidates. Interestingly the stability of the Majoron as dark matter is related to the scale that sets the see-saw and leptogenesis mechanisms, while its annihilation and scattering cross section off nuclei can be set through the Higgs portal. For $\mathcal{O}(GeV) - \mathcal{O}(TeV)$ Majorons, we compute observables such as the abundance, scattering cross section, Higgs invisible decay width, and emission lines and compare with current data in order to outline the excluded versus still viable parameter space regions. We conclude that the simplest Majoron dark matter models coupling through the Higgs portal, except at the Higgs resonance, are excluded by current direct detection data for Majorons lighter than $225$~GeV and future runnings are expected to rule out decisively the 1GeV-1TeV window. Lastly, we point out that light keV-scale Majorons whose relic density is set by thermal freeze-in from sterile neutrinos can account for the keV line observed by XMM-Newton observatory in the spectrum of 73 galaxy clusters, within a see-saw model with a triplet Higgs. | The identity of dark matter constitutes one of the most exciting puzzles in current science. Interestingly, dark matter is often connected to other paradigms of fundamental physics, with WIMPs (Weakly Interacting Massive Particles) in supersymmetric theories being the most studied ones. The dark sector can also be connected to other important phenomena such as the generation of the neutrino masses, leptogenesis, and baryogenesis \cite{Tsuyuki:2014aia,Kashiwase:2013uy,seesaw}. The Majoron dark matter model is an example which occurs in see-saw models of neutrino mass generation. In see-saw models, the lepton number might be explicitly broken by the Majorana masses of right-handed neutrinos. If the lepton number is instead broken spontaneously by the vev of a complex scalar field (a singlet ``Higgs"), one has a new pseudo-scalar gauge singlet Nambu-Goldstone boson (the Majoron). In such models, the Majoron is a natural decaying dark matter candidate \cite{ernestma,CMBmajoron,Gelmini:1984pe,Berezinsky:1993fm}. The Majoron lifetime is determined by its decay into Standard Model (SM) neutrinos which is suppressed by the scale of lepton number violation. For sufficiently high lepton violation scale the Majoron is cosmologically stable. This scale also sets the heavy right handed neutrinos masses in the see-saw type I setup. Therefore, the viability of the Majoron dark matter is connected to the see-saw mechanism responsible for generating the SM neutrinos masses and the scale of leptogenesis which occurs through the decays of the heavy right-handed neutrinos \cite{Covi:2009xn}. As for the mass of the Majoron, it can arise from explicit soft terms, or from quantum gravitational effects that explicitly break lepton number. In the former case, the mass can be hundreds of GeV. From an effective field theory point of view, nothing prevents a coupling of the Majoron to the Higgs scalar potential at tree level. Majoron models of this category are thus a particular UV realization of the effective Higgs-portal scalar models studied by Ref.\cite{Burgess:2000yq}. In the case where the mass is due to quantum gravitational effects, on the other hand, the Majoron is expected to be very light. A particularly well-motivated scenario is a $\mathcal{O}($keV$)$ Majoron, which can satisfy the thermal relic density. We study dark matter observables in both cases. In the former, we compute % observables such as the abundance, scattering cross section, Higgs invisible decay width, and emission lines and compare with current data in order to outline the excluded versus still viable parameter space regions in the $\mathcal{O}(GeV) - \mathcal{O}(TeV)$ mass range. We find that LUX bounds on the dark matter scattering cross section, along with relic density requirements and Higgs invisible decay width limit, effectively rule out thermal dark matter below $\sim 225$ GeV in the this scenario (except near the Higgs resonance when the Majoron mass is $\sim 60$ GeV). Furthermore, future direct detection running coming from XENON1T and LUX are forecasted to rule out the entire GeV-TeV mass range. A way out would be the inclusion of new particles with masses close to the Majoron in order to exploit co-annihilation channels. Additionally, a non-thermal history for the Majoron, where the initial number density of Majorons is fixed without further annihilation, is possible \cite{Allahverdi:2010rh}. In the case of a $\mathcal{O}(keV)$ Majoron, we point out that it may be possible to accommodate the recently observed keV line from the XMM-Newton observatory \cite{KeVline}, for appropriate choice of parameters in a see-saw model with a triplet (as well as the singlet) Higgs scalar. The branching to photons arises at loop level from the projection of the Majoron along the doublet (Standard Model) Higgs. The relic density can be set by thermal freeze-in from the right-handed neutrinos. We provide order-of-magnitude estimates as a proof of concept that the observed keV line can be obtained in this class of models. The plan of the paper is as follows. In Section \ref{model}, we describe the Majoron dark matter model in the context of type I see-saw. In Section \ref{darkmatter}, we discuss the dark matter observables of the model. In Section \ref{kevmajoron}, we discuss the case of a light Majoron and the recently observed keV line. We end with our Conclusions. | Majorons are Nambu-Goldstone bosons arising from the spontaneous breaking of lepton number symmetry by a complex scalar. The Majoron mass is model-dependent, although it is expected to be small, due to explicit lepton symmetry breaking by quantum gravity effects. On the other hand, explicit soft terms can be introduced to make the Majoron mass large. These scalar dark matter models have several interesting features: $(1)$ they are examples of decaying dark matter, with the decay being mainly to neutrinos, suppressed by the scale of lepton symmetry breaking. $(2)$ in the heavy (GeV - TeV) Majoron case, this model is a UV realization of the effective Higgs-portal scalar dark matter framework. With this in mind we have computed the Higgs invisible decay width into Majorons, abundance, direct and indirect detection observables and compared them with the most current data available. The Majoron has a somewhat large parameter space that can reproduce the right abundance as shown in Fig.1. In case the Majoron is sitting at the Higgs resonance the model is consistent with all current bounds. Otherwise the recent constraint on the scattering cross section reported by LUX in 2013 decisively rules out the thermal Majoron window for $M_J < 225$~GeV in the simplest Majoron model which has a singlet scalar and heavy-right handed neutrinos. Additionally, future direct detection running coming from XENON1T and LUX are expected to rule out the entire GeV-TeV mass range. Therefore either non-thermal production mechanisms or the inclusion of new particle to exploit co-annihilation channels are required to circumvent this conclusion. Another alternative would be the inclusion of new particles to play to role of the mediator in new annihilations channels. We have also discussed indirect detections bounds. In particular the late decay of Majorons may distort the CMB power spectrum, therefore a bound of $\Gamma < 10^{-19} s^{-1}$ is required. This enforces the Majoron mass to not be much greater than the TeV scale for a large scale where the lepton number is softly broken. If one pushes down the latter scale the upper bound on Majoron is rapidly strengthened according to Eq.16. $(3)$ lastly, we discussed a Majoron model which has a Higgs triplet in its spectrum in light of the recent 3.5~keV line and concluded that the required signal could be obtained for a lepton number violation scale $\sim \mathcal{O}(10^4)$ TeV. We noted that the thermal freeze-out scenario does not address such a line because the mass of the Majoron cannot be larger than $\sim 0.1$~keV. However, if one uses thermal freeze-in through the sterile neutrino portal to set the Majoron relic density (alternatively, a non-thermal mechanism may also work), then it is possible to have the Majoron as the possible candidate to this $3.5$~keV signal while being consistent with other x-ray searches according to Fig.2. We provided a proof of concept that the observed keV line can be obtained in this class of models, for plausible choices of the model parameters. The relation between light decaying dark matter, line emission bounds, and thermal freeze-in certainly warrants further study \cite{farinaldokuver}. | 14 | 4 | 1404.1400 |
1404 | 1404.3858_arXiv.txt | Hesperian chaotic terrains have been recognized as the source of outflow channels formed by catastrophic outflows. Four main scenarios have been proposed for the formation of chaotic terrains that involve different amounts of water and single or multiple outflow events. Here, we test these scenarios with morphological and structural analyses of imagery and elevation data for Aram Chaos in conjunction with numerical modeling of the morphological evolution of the catastrophic carving of the outflow valley. The morphological and geological analyses of Aram Chaos suggest large-scale collapse and subsidence (1500 m) of the entire area, which is consistent with a massive expulsion of liquid water from the subsurface in one single event. The combined observations suggest a complex process starting with the outflow of water from two small channels, followed by continuous groundwater sapping and headward erosion and ending with a catastrophic lake rim collapse and carving of the Aram Valley, which is synchronous with the 2.5 Ga stage of the Ares Vallis formation. The water volume and formative time scale required to carve the Aram channels indicate that a single, rapid (maximum tens of days) and catastrophic (flood volume of 9.3$\cdot$10$^4$ km$^3$) event carved the outflow channel. We conclude that a sub-ice lake collapse model can best explain the features of the Aram Chaos – Valley system as well as the time scale required for its formation. | \label{Introduction} Martian chaotic terrains are deeply collapsed areas ($>$ 1 km deep) that stretch for up to hundreds of kilometers and show a bumpy floor characterized by an irregular pattern of fractures and tilted blocks of different sizes (from meter to kilometer scale). Chaotic terrains predominantly occur along the dichotomy boundary between the southern highlands and northern lowlands \citep{Sharp1973,Chapman2002,Rodriguez2005a,Glotch2005,Meresse2008,Warner2011}. Outflow channels represent the largest systems carved by liquid water on Mars. They are thousands of kilometers long, more than a kilometer deep \citep{Baker2001}and show attributes such as grooves, terraces, teardrop islands, streamlined terraces and high width-to-depth ratios that are consistent with the erosive origin of the channels \citep[e.g.,][]{Nelson1999,Baker2001,Coleman2005,Pacifici2009,Warner2010a}. In many cases, chaotic terrains represent the source area of Hesperian (approximately 3.7--3.3 Ga) outflow channels \citep{Nelson1999,Tanaka2003} and several authors \citep[e.g.,][]{Carr1979,Carr1996a,Baker2001} argue that those chaotic terrains were formed by a rapid discharge of water from the subsurface, resulting in collapsed and fractured areas and massive flows carving the large outflow channels. However, the actual evolutionary process leading to chaotic terrain formation and collection and the discharge of catastrophic volumes of water ($\ge$10$^5$ km$^3$) has remained controversial. Several evolutionary processes have been proposed that can be grouped in four different scenarios (Fig. 1). In the first hypothesis (Fig. 1a), fracturing in the bedrock led the water expulsion to the surface. The water was generated by partial melting of the cryosphere after magmatic intrusions that increased the subsurface temperature \citep[e.g.,][]{Sharp1973,Chapman2002,Ogawa2003,Rodriguez2005a,Leask2006b,Meresse2008}. \begin{figure*}[ht!] \centerline{\includegraphics[width=.75\textwidth]{figure1.pdf}} \caption{Four different scenarios for the origin of chaotic terrains. a) Magma-cryosphere model \citep[redrawn after][]{Meresse2008}; b) Aquifer model \citep[redrawn after][]{Harrison2009}; c) Gas-hydrate model \citep[redrawn after][]{Kargel2007}; d) sub-ice lake model \citep[redrawn after][]{Zegers2010}.} \label{map} \end{figure*} In a second group of mechanisms (Fig. 1b), the release of water from the cryosphere was the result of an increase in pressure of a global pressurized sub-cryospheric aquifer \citep[e.g.,][]{Carr1979,Clifford1993,Hanna2007a,Harrison2009}. The discharge from the pore space relied on the flow of water through a permeable subsurface layer, where water that was discharged was replaced by recharge. In the majority of martian aquifer models, water is generally assumed to have recharged from a great distance \citep[$\ge$2000 km,][]{Clifford1993,Harrison2008}. Numerical flow models \citep{Hanna2007a,Harrison2008} indicate that the large total volume of liquid water (10$^5$--10$^6$ km$^3$) and high discharge rates (10$^6$--10$^9$ m$^3$/s) required to form the morphology of the outflow channels were not achievable by a single discharge from a porous medium. Because of the flow volume discharge from aquifers, a scenario was proposed wherein a large number of small flooding events were followed by a sudden release of previously ponded water \citep[$\ge$600,][]{Harrison2008}. Another hypothesis (Fig. 1c) suggests that dewatering of fluids from gas and/or salt hydrate buried deposits \citep{Max2001,Montgomery2005} or the hydrologic processes triggered by clathrates \citep{Kargel2007} could be responsible for the water outflow. Finally, \citet{Zegers2010} propose that chaotic terrains developed by the catastrophic collapse of sediments induced as a consequence of the melting of buried ice sheets (Fig. 1d). Thermal modeling results show that even under very low crustal heat flux conditions, ice sheets will melt if buried under thick sediments (up to 2 km) because of the difference in thermal conductivity between the basin fill and surrounding crust. When the buried sub-ice lake reaches a critical thickness, the overburden collapses and subsides, resulting in a massive expulsion of water to the surface. To distinguish between those four scenarios and test the validity of the evolution models for chaotic terrains and their outflow channels, the amount and timing of the water release, the amount of subsidence, and the fracture distributions are fundamental variables. To constrain those variables for a single chaotic terrain and its outflow channel, we analyzed the morphological and geological features characterizing Aram Chaos and its valley. Furthermore we estimated the flow volume and formative time scale required to carve the Aram channel. Finally we discuss the four scenarios proposed for chaotic terrain formation in view of these results, propose the most likely scenario for the evolution of Aram Chaos and Aram channel and discuss the significance of our findings for other chaotic terrains. | The geological and hydrological analyses performed on the Aram Chaos-Valley system indicate that the flow volume required to carve the Aram Valley and two small channels (9.3$\cdot$10$^4$ km$^3$) is similar to the volume of water that could have been produced in an event of Aram Chaos by the melting of a buried ice lake \citep[9.2$\cdot$10$^4$ km$^3$,][]{Zegers2010}. The formative time evaluation confirms that a single, rapid (tens of days) and catastrophic event was sufficient to carve the channel rather than many small groundwater events active for a relatively long time. The resulting Aram Chaos morphology implies large amount of subsidence (1500 m) and one (or two) intense fracturing event(s). Such a scenario is consistent with a model of a buried sub-ice lake system \citep{Zegers2010}. The thermal insulation and relatively low heat flow may have been sufficient to induce ice melting. When the system became unstable, a massive water outflow occurred with the collapse of sediment and carving of the outflow channel. The sub-ice lake scenario explains many features characteristic of chaotic terrains, although more than one mechanism may have been involved in the carving of larger outflow channels. | 14 | 4 | 1404.3858 |
1404 | 1404.6199_arXiv.txt | {Turbulent transport of chemical elements in radiative zones of stars is considered in current stellar evolution codes thanks to phenomenologically derived diffusion coefficients. Recent local numerical simulations (Prat \& Ligni\`eres 2013, \aap, 551, L3) suggest that the coefficient for radial turbulent diffusion due to radial differential rotation satisfies $D_{\rm t}\simeq0.058\kappa/Ri$, in qualitative agreement with the model of Zahn (1992, \aap, 265, 115). However, this model does not apply (i) when differential rotation is strong with respect to stable thermal stratification or (ii) when chemical stratification has a significant dynamical effect, a situation encountered at the outer boundary of nuclear-burning convective cores.} {We extend our numerical study to consider the effects of chemical stratification and of strong shear, and compare the results with prescriptions used in stellar evolution codes.} {We performed local, direct numerical simulations of stably stratified, homogeneous, sheared turbulence in the Boussinesq approximation. The regime of high thermal diffusivities, typical of stellar radiative zones, is reached thanks to the so-called small-P\'eclet-number approximation, which is an asymptotic development of the Boussinesq equations in this regime. The dependence of the diffusion coefficient on chemical stratification was explored in this approximation.} {Maeder's extension of Zahn's model in the strong-shear regime (Maeder 1995, \aap, 299, 84) is not supported by our results, which are better described by a model found in the geophysical literature. As regards the effect of chemical stratification, our quantitative estimate of the diffusion coefficient as a function of the mean gradient of mean molecular weight leads to the formula $D_{\rm t}\simeq 0.45\kappa(0.12-Ri_\mu)/Ri$, which is compatible in the weak-shear regime with the model of Maeder \& Meynet (1996, \aap, 313, 140) but not with Maeder's (1997, \aap, 321, 134).} {} | Mixing of chemical elements in stellar interiors is a crucial feature in stellar evolution theory. Indeed, for main sequence stars, mixing can continuously draw hydrogen from outer layers to the core. For a given initial mass, stars experiencing mixing have a higher averaged mean molecular weight and thus appear brighter. The hydrogen-core-burning lifetime is increased as is the mass of the helium core. Larger helium cores modify the subsequent evolution. It is thus essential to have a good knowledge of transport processes to build reliable stellar models. Whereas microscopic processes, such as molecular diffusion, gravitational settling, and radiative acceleration, are relatively well known, the effects of macroscopic motions induced by rotation are still poorly understood. Improving such models may have a strong impact in many fields of astrophysics. This is particularly true for galaxy physics, because the properties of the stars in a galaxy provide information about the galaxy, and for planetary systems, where constraints on the stellar host help us constrain the mass and the radius of planets. The main technique used to constrain mixing processes in stars is to determine chemical abundances on the surface of stars through measurements of the depth and the width of atomic absorption lines. These measurements may bring out some anomalies for certain stars, that is, over- or under-abundances as compared to the standard model of stellar evolution, which describes the evolution of a spherical, non-rotating star in which the only macroscopic transport process is convection. For massive stars, in which the CNO cycle is dominant, abundance anomalies of elements involved in this cycle, such as helium, carbon, or nitrogen, indicate that some deep mixing is occurring in the observed star (see e.g. \citealt{GiesLambert}, \citealt{Lyubimkov}, and more recently \citealt{Hunter} with the VLT-FLAMES survey). \citet{Martins2009, Martins2013} have even found evidence for the existence of stars in which mixing is so efficient that they are quasi-homogeneous. Another sort of anomaly is linked to the fact that light elements (lithium, beryllium, and boron) are destroyed at high temperature. For massive stars, lithium and beryllium are totally depleted, whereas boron is only destroyed in the inner layers. If some mixing occurs between these layers and the outer ones, the surface abundance of boron is lower than otherwise \citep{Venn, Mendel}. For cooler stars, deep convection causes lithium and beryllium to be depleted. It follows that these elements can only be detected in a narrow range of stellar mass. Here again we find discrepancies between the standard model and the observed abundances, including for the Sun \citep[e.g. the lithium dip, initially observed by][]{BoesgaardTripicco}. A source of extra mixing, such as internal gravity waves \citep{TalonCharbonnel}, is thus needed. Stellar seismology is another powerful way to constrain transport processes within stars. The analysis of stellar oscillations not only provides new global parameters that are useful for better determining fundamental parameters of stars, such as mass, radius, and age, but it also allows us to probe the internal properties of some stars, including the distribution of chemicals (for the Sun), \emph{via} the speed of sound \citep{ChristensenDalsgaard, AntiaBasu} and internal rotation. \citet{Beck} and \citet{Mosser} have notably determined the rotation rate in the core of red giants, and a rotation profile has even been obtained for the Sun \citep{Thompson} and a subgiant \citep{Deheuvels}. Thanks to this promising technique, we are about to be given access both to the causes of the transport (rotation) and to its effects (chemical distribution). We still need to extend its use to a larger number of stars, which is in progress, especially for CoRoT and Kepler stars. In many stellar evolution codes, rotationally induced mixing is taken into account thanks to a set of turbulent diffusion coefficients linked to various hydrodynamical instabilities triggered by differential rotation and meridional circulation. Among these instabilities, shear instability is thought to produce the most efficient mixing \citep{KnoblochSpruit}, sometimes called ``shear mixing''. This instability normally occurs in a stably stratified sheared flow when the Richardson number $Ri=(N/S)^2$, which compares stable stratification (with the Brunt-Väisälä frequency $N$) and the shear (with the shear rate $S={\rm d}U/{\rm d}z$), is lower than a critical value $Ri_{\rm c}$, which equals $1/4$ in the inviscid linear stability analysis \citep{Miles}. However, except in some limited layers of evolved stars \citep{Hirschi}, the Richardson number is generally much larger than one in stars, and this criterion would result in almost always shear-stable stellar interiors. In this high-Richardson-number or weak-shear regime, the very high thermal diffusivities found in stars are nevertheless able to reduce the amplitude of the buoyancy force through radiative losses and thus to overcome the stabilising effect of stratification \citep{Townsend, Zahn1974}. We are interested here in one of the transport coefficients related to this instability, the radial diffusion coefficient due to radial differential rotation. It was initially derived by \citet{Zahn1992} from the following phenomenological arguments. \begin{itemize} \item Turbulent flows generally tend to reach a statistical steady state that is marginally stable with respect to the instability that has generated them. \item The destabilising effect of the very high thermal diffusivity $\kappa$ is such that the Richardson instability criterion should be replaced by $RiPe < Ri_{\rm c}$, where $Pe=UL/\kappa$ denotes the P\'eclet number based on velocity and length scales $U$ and $L$. \item The relevant P\'eclet number to use in this criterion is the turbulent one $Pe_\ell=u\ell/\kappa$, where $u$ is the velocity scale and $\ell$ the length scale of turbulence. \item The turbulent diffusion coefficient is proportional to the product of these turbulent scales: $D_{\rm t}\simeq u\ell/3$. \end{itemize} Zahn finally obtains a diffusion coefficient that reads \begin{equation} D_{\rm t}\simeq\frac{\kappa}{3}\frac{Ri_{\rm c}}{Ri}. \end{equation} Since, several attemps have been made to add various physical ingredients to this model, such as $\mu$-gradients \citep{MaederMeynet1996, Maeder1997} and horizontal diffusion \citep{TalonZahn}. As regards massive stars, the use of such models of rotational mixing in stellar evolution codes gives results closer to the observations than the standard model both for CNO products \citep{HegerLanger, MeynetMaeder2000, MaederMeynet2001} and for boron \citep{Fliegner, Mendel, Frischknecht}. Thanks to the progress in the computational efficiency of such codes, it is now possible to compute entire synthetic stellar populations from a given distribution of initial conditions that include chemical composition and angular momentum. As shown by \citet{Brott} and \citet{Potter}, these synthetic populations can be compared to observed ones, such as those obtained by large surveys. In particular, the authors recover the main group of stars observed by \citet{Hunter}. However, some observational constraints are still not explained by these models. In particular, \citet{Hunter} are unable to explain the existence of nitrogen-enriched slow rotators which are actually observed. Besides, current models do not predict enough transport of angular momentum to recover the seismic constraints on internal rotation in red giants and subgiants \citep{Eggenberger, Ceillier, Marques}. Most of these phenomenological models of turbulent transport have never really been tested. Indeed, the physical conditions present in stellar interiors are far too extreme to allow us to perform realistic laboratory experiments. Our approach is to use direct numerical simulations (DNS) as experiments to test the existing prescriptions. Local numerical simulations of homogeneous, sheared, and stably stratified turbulence (such as those performed in a geophysical context by \citealt{Gerz}, \citealt{Holt}, and \citealt{Jacobitz}) have the advantage of being as generic as possible while taking the essential ingredients (shear and stable stratification) into account. There has already been an attempt to test Zahn's model by \citet{BruggenHillebrandt} but because their simulations used numerical viscosity and thermal diffusivity, which depend on the resolution, they were unable to study the effect of thermal diffusion. In a previous paper \citep[hereafter Paper I]{PratLignieres}, we explored the dependence of the radial turbulent diffusion coefficient on thermal diffusity in the regime of the small P\'eclet numbers typical of stellar interiors. This was done by considering a parallel-plane flow configuration with forced, uniform, vertical stable stratification and velocity shear. Boundary conditions were periodic in the horizontal directions and impermeability, along with the mean shear, were imposed at the lower and upper boundaries. Since the very high thermal diffusivity introduces a huge gap between the diffusive time scale and the dynamical one, which drastically increases the computational cost of complete Boussinesq simulations, we used an asymptotic development of the Boussinesq equations called the small-P\'eclet-number approximation (SPNA) described by \citet{Lignieres}. We found that our simulations are qualitatively in good agreement with Zahn's prescription in this regime and we were able to give a quantitative estimate of the turbulent diffusion coefficient. The purpose of the present paper is twofold. On the one hand, we extend the previous study to the domain of large P\'eclet numbers, where Zahn's model is not valid any more. Although the net effect on global mixing is often assumed to be negligible, some strong-shear situations exist, especially in evolved stars \citep[see for example][]{Hirschi}, where the Richardson number is small enough that stellar interiors are unstable with respect to the shear instability even on large scales, thus at large P\'eclet numbers. On the other hand, we explore the dependence of the turbulent diffusion coefficient with respect to chemical stratification. This effect is likely to be significant at the frontier of the radiative zone and a convective core. Indeed, strong chemical gradients are expected in such a boundary, so that whether mixing occurs or not has a strong influence on stellar evolution. Section~\ref{sec:form} presents the formalism and the methods used in our simulations. Section~\ref{sec:large} is devoted to the case of a neutral chemical stratification, whereas Sect.~\ref{sec:chim} deals with the case of stable chemical stratification. Then, the astrophysical consequences of our results are discussed in Sect.~\ref{sec:disc}, along with some prospects. | \label{sec:disc} We have tested several models of turbulent transport. For large P\'eclet numbers, our simulations invalidate the model of \citet{Maeder1995} and validate that of \citet{LindborgBrethouwer}. Nevertheless, the latter cannot be applied as such in stellar evolution codes. As regards the dynamical effect of chemical stratification, our simulations agree with a special case of the model of \citet{MaederMeynet1996}, but not with that of \citet{Maeder1997}, which nevertheless gives the better results when compared with observations. These results were obtained with a flow configuration that depends on several physical and numerical parameters, such as initial conditions and the Reynolds number. As regards initial conditions, we have chosen a parameter domain in which the flow is likely to be the most generic, as discussed in Sect.~\ref{sec:init}. For the Reynolds number, a preliminary study shows that transport does not significantly vary for turbulent Reynolds numbers ranging from 225 to 340. According to \citet{MichaudZahn}, these values are consistent with what we expect in stellar radiative zones. In addition, this configuration relies on certain hypotheses, including stationarity and homogeneity, that are forced in an uncommon way. This configuration is generic, but not necessarily natural. Other numerical methods such as shearing boxes, used notably in geophysics \citep[e.g.][]{Jacobitz} and in astrophysics for the study of accretion discs \citep[e.g.][]{Lesur}, also force uniform mean shear flows, but not exactly in the same way as in the present paper. It would be interesting to see if our results can be recovered with such approaches. Moreover, unstationary and unhomogeneous configurations should also be considered. As an example, direct numerical simulations of free shear layers, such as in \citet{BruggenHillebrandt}, could be considered. There are still many other physical ingredients that could play a role in the radial transport due to shear instability. Among them is the effect of a very efficient horizontal turbulent diffusion, which is said to reduce the stabilising effect of chemical stratification \citep{TalonZahn}. One can also wonder what the effects of rotation (by introducing the Coriolis force in the simulations) and magnetic field are on shear instability. An attempt has been made recently by \citet{Maeder2013} to provide some general diffusion coefficients, when simultaneously taking several hydrodynamical instabilities into account. As regards magnetic field, \citet{Lecoanet} show that in the linear regime particular magnetic configurations can destabilise a flow, which would be stable without it. Whether this result remains relevant in the non-linear regime is still an open question. Until now, we have computed transport coefficients for chemical elements, but turbulent viscosity is also crucial to understanding angular momentum transport and should be estimated in the same way. In a forthcoming paper, we intend to determine relations between turbulent viscosity and turbulent thermal and chemical diffusivities. | 14 | 4 | 1404.6199 |
1404 | 1404.6961_arXiv.txt | {The scientific mission of ESTCube-1, launched in May 2013, is to measure the Electric solar wind sail (E-sail) force in orbit. The experiment is planned to push forward the development of E-sail, a propulsion method recently invented at the Finnish Meteorological Institute. E-sail is based on extracting momentum from the solar wind plasma flow by using long thin electrically charged tethers. ESTCube-1 is equipped with one such tether, together with hardware capable of deploying and charging it. At the orbital altitude of ESTCube-1 (660--680~km) there is no solar wind present. Instead, ESTCube-1 shall observe the interaction between the charged tether and the ionospheric plasma. The ESTCube-1 payload uses a 10-meter, partly two-filament E-sail tether and a motorized reel on which it is stored. The tether shall be deployed from a spinning satellite with the help of centrifugal force. An additional mass is added at the tip of the tether to assist with the deployment. During E-sail experiment the tether shall be charged to 500~V potential. Both positive and negative voltages shall be experimented with. The voltage is provided by a dedicated high voltage source and delivered to the tether through a slip ring contact. When the negative voltage is applied to the tether, the satellite body is expected to attract electron flow capable of compensating for the ion flow, which runs to the tether from the surrounding plasma. With the positive voltage applied, onboard cold cathode electron guns are used to remove excess electrons to maintain the positive voltage of the tether. In this paper we present the design and structure of the tether payload of ESTCube-1.} | The scientific mission of ESTCube-1, the first Estonian satellite, is to perform an on-orbit test of the Electric solar wind sail (E-sail) concept. The E-sail is a propulsion innovation made at the Finnish Meteorological Institute (FMI) in 2006. \cite{Janhunen2007AnnGeophys} Thrust is produced by harnessing the momentum of charged solar wind particles by using long, thin, and electrically charged conducting tethers. A fullscale E-sail spacecraft is planned to have e.g. 100 tethers, each 20 km long. The sail is kept open with the help of centrifugal force, the spacecraft must therefore be kept in rotating motion. Once operational, E-sail technology is expected to revolutionize the space travel within our solar system. \cite{JanhunenProcEstAcadSci2014, Janhunen2008JBIS, Mengali2009CelMechDynAstron, Janhunen2009AnnGeophysNegative, Toivanen2009, Quarta2010HazardAsteroids, Janhunen2010RevSci, Merikallio2010Asteroidit} The experiment conducted by ESTCube-1 marks the beginning of space test era for the E-sail. \cite{Silver2014} The ESTCube-1 payload (PL) includes one E-sail tether, deployable to $\sim$10-meter length. The tether is stored on a motorized reel and it can be reeled out upon request from the ground. The principle of tether deployment is similar to the proposed principle of larger E-sails, i.e. the satellite is spun around its axis of maximum moment of inertia and the exiting tether is streched by the effect of centrifugal force. This force is enhanced by placing a small end mass at the tip of the tether. Successful deployment of the tether is verified by observing a noticeable drop in the satellite spin rate. In addition, a visual verification shall be obtained by imaging the tether and its end mass with the onboard camera during the tether deployment. Once deployed, the tether shall be charged with a voltage of 500 V. Both positive and negative voltages, corresponding to positively and negatively charged E-sails, respectively, shall be tested. Because of its orbit (a low-Earth orbit with 660--680 km altitude) ESTCube-1 is not influenced by solar wind. Instead, the charged tether shall interact with the ionospheric plasma, through which the satellite is traveling with its orbital speed of 7.5 km/s. When the on and off cycles of the tether voltage are correctly synchronized to the satellite's spin rate, the E-sail effect between the tether and the plasma can be observed as a cumulative change in the spin rate. The change can be chosen to be in either direction, depending on the on/off cycling. In this paper we give an overall technical description of the onboard apparatus which shall be used to perform the E-sail experiments. The satellite was launched into orbit on May 7th 2013 (UTC), but at the time of writing it had not yet entered the tether experiment phase of its mission. | This paper has described the contents of the ESTCube-1 nanosatellite payload. The payload is used for carrying out the first space experiment of the E-sail concept. The equipment was built and delivered to the ESTCube-1 nanosatellite team in time. The satellite was integrated to the European Space Agency's (ESA) Vega launch vehicle and the satellite was launched and delivered into orbit on May 7th, 2013 (UTC) from Europe's Spaceport in Kourou, French Guiana. The first vibration tests revealed a design flaw in the motor board. The piezo rotator was sensitive to certain resonance frequencies of the vibrations. This caused the tether reel to turn and thus break the tether. This flaw was subsequently fixed by introducing the reel lock. The reel lock itself went through some vibration testing, but the whole tether reel system did not. However, at this point we already know almost with full certainty that the reel lock has accomplished its locking duties. A reel lock failure, or a tether breakage of any sort, would almost certainly lead to the loose tether filling the insides of the satellite, creating short cuts at several places. During its time in orbit the satellite has mostly functioned as expected. This does not support the notion of any launch failures having taken place. Also, the electron guns were not tested for vibrations. This was a planned strategy. Because of the very early stage of development of the miniaturized cold cathode electron gun technology, only flight models were available within the schedule of our mission. However, unlike the tether and its related equipment, the electron guns are not mission critical instruments. Even with no functional electron guns the negative voltage mode can still be used for the E-sail experiment. In addition to the matters discussed above, no particular adversities have emerged. The systems worked flawlessly in functional ground tests. At the time of the writing (June 2013) the project team in Tartu is validating the operations of the satellite in orbital conditions. Also the final flight software is being written. Schedule for the E-sail tether experiment has not yet been fixed. Once made, the experiment is also expected to provide some useful data for our next mission. In 2014 or 2015 it is time to launch the first Finnish satellite, Aalto-1. It is a 3U CubeSat designed and built by the students of Aalto University, Espoo, Finland. A similar tether payload will be included in Aalto-1 as one of its secondary payloads. The length of the tether in Aalto-1 will be 100 meters. {\it Acknowledgements\ }Following persons are acknowledged for their contribution to this work. Jouni Polkko and Sini Merikallio from FMI. Risto Kurppa, Tuomo Ylitalo and G\"oran Maconi from UH. Olaf Kr\"omer and Roland Rosta from DLR. Pekka Salminen from SkyTron co. Alexander Obraztsov from the University of Eastern Finland Joensuu. | 14 | 4 | 1404.6961 |
1404 | 1404.6685_arXiv.txt | We estimate the quasiclassical probability of the homogeneous nuclear matter transition to a strange matter when a detonation wave propagates radially inside a sphere of nuclear matter. For this purpose we make use of instanton method which is known in the quantum field theory.\\ {\bf Keywords:} Strange matter, Instanton. {\bf PACS 2006:} Primary -- 47.40.Rs, 03.65.Xp, Secondary -- 26.60.-c, 97.60.Jd | \label{s1} It was first pointed out \cite{bo71,wi84} that strange matter (SM) composed of three quarks might be a ground state of a normal nuclear matter (NM) at zero temperature and pressure, which was later supported by studies based on MIT bag model \cite{FJ84}. A conversion of NM to SM is suppressed at ordinary nuclear densities. The existence of stable SM would have some remarkable consequences in cosmology and astrophysics. At very large densities of NM like those in neutron stars (NS), where the Fermi energy is higher than mass of {\em s} quark, the NM--SM transition may occur spontaneously. This led to conjecture \cite{al88, bo71,ms88,wi84} that strange stars, which are predominantly made of SM, may be formed from dense NS. The conversion is assumed to be triggered at the core of NS \cite{al86,ol87} where the density reaches values $2\cdot 10^{14}g/cm^3\!<\! \rho_*\!<\!6\cdot 10^{15}g/cm^3$ with a total mass of the star $M\geq 1.5\odot$. There may appear stable SM drops, called {\em strangelets} \cite{FJ84}, if every single drop possesses a baryon charge $A$ exceeding some critical value $A_*$. Further growth of strangelets occur by outward diffusion of strange quarks to ambient NM \cite{al86,ol87}. Equation of SM state has been suggested in \cite{wi84}, $P_s=(E_s-E_o)/3$, where $P_s$ and $E_s$ stand for the pressure and density of energy, and $E_o$ denotes a density of energy of SM at zero pressure. If $E_s\gg E_o$ then transition from the non relativistic NM ($P_n\ll E_n$) to SM occurs with essential growth of pressure and temperature. There are two different models which treat the NM--SM transition in framework of relativistic hydrodynamics: combustion waves (CW) \cite{al86, ol87} and detonation waves (DW) \cite{ho88, to05}. The CW propagates as a slow combustion with a speed $V_c\simeq 10^7m/s$, while the DW propagates with $V_d\simeq 10^8 m/s$. In \cite{to05} DW was considered as the self-similar spherical wave propagating with a constant rate w.r.t. NM of constant density. Different aspects of this conversion were discussed in \cite{ho88, ma11}. The problem arises when the classical solution is considered at the strangelet scale with radius $R_s=(3Am_n/4\pi\rho)^{1/3}$, where $\rho$ denotes a density of NM and $m_n$ stands for neutron mass. For strangelets with baryon charge $A \simeq 10-100$ this radius varies in the range $R_s\simeq 1.2-2.5\cdot 10^{-15} m$. On the other hand, the de Broglie wavelength $\lambda_B=h/(Am_nV_d)$ for the strangelet reads, $\lambda_B\simeq 0.4-4\cdot 10^{-15}m$, {\em i.e.}, both $\lambda_B$ and $R_s$ have comparable values. This manifests the quasiclassical nature of the strangelets which trigger the NM--SM transition and poses a question about probability of such transition. To answer this question we make use of the known in quantum theory instanton approach \cite{ra82} which describes a tunneling between different field configurations. We calculate a probability of the NM--SM conversion when DW propagates spherically inside NM. | \label{s5} In the framework of instanton approach we have shown that NS with the core density $\rho_*\simeq 10^{15}g/cm^3$ allows to have at least one stable strangelet during the time star existence $T_2$, $T_N<T_2<T_U$, if the baryon number is $A_*=24$, where $T_N\simeq 10^6$ years and $T_U\simeq 13.8\cdot 10^9$ years stand for the largest time NS existence and the universe age, respectively. A low value of $A_*$ makes it interesting to compare it with those discussed in literature. For $2\!<\!A\!<\!6$ quantum chromodynamics strongly suggests complete instability of any strangelets \cite{jw99}. In \cite{FJ84} the SM is studied for low $A<\!10^2$ and large $10^2\!<\!A\!<\!10^7$ baryon numbers. This wide range covers many other values for $A$ discussed in literature: $A\simeq 16-40$ \cite{bo71}, $A<10^2$ \cite{gj93}, $A>10^2$ \cite{ol87}, $A\simeq 10^3$ \cite{wi84}, $A\simeq 10^2-10^4$ \cite{ms02} and most of these values are substantially larger than $A_*$. We put forward an agent which may be responsible for the higher $A_*$ in the framework of NM--SM instanton transition. The mass of equilibrium configuration of cold matter at each central star density $\rho_*$ ($g/cm^3$) is a damped periodic function of $\ln\rho_*$ \cite{ha65,ze96}. There are two ranges for which these configurations are stable : the white dwarfs with low electron density, $10^5<\!\rho_*<\!10^8$, and the neutron stars with high density, $10^{14}\!<\!\rho_*\!<\!\rho^{OV}$ where $\rho^{OV}\!\simeq\! 6\cdot 10^{15}$ denotes the Oppenheimer-Volkoff limit. There are also a number of extrema for $\rho_*$ exceeding $\rho^{OV}$: such superdense configurations were found in \cite{mz64}, $10^{18}\!<\!\rho_*\!<\! 10^{20}$, and in \cite{ha65}, $\rho_*>3\cdot 10^{21}$. In Figure \ref{fg3} we show how the admissible values of $A$ do increase once the density of the NS core grows, e.g., for $\rho_*\simeq 10^{19}$ we have $240<A_*<250$. \begin{figure}[h!]\begin{center} \psfig{figure=./Inst4.eps,height=5cm,width=14cm} \end{center} \vspace{-.7cm} \caption{Plots of the functions $t(A)$ defined in (\ref{k22}) for different densities $\rho_*$ ($g/cm^3$) in the NS core: $10^{15}$ ({\em red}), $10^{16}$ ({\em blue}), $10^{17}$ ({\em green}), $10^{18}$ ({\em magenta}), $10^{19}$ ({\em brown}), $10^{20}$ ({\em black}). Two {\em dashed lines} mark two time scales $T_N=10^{t_n}$, $t_n\simeq 13.5$, and $T_U=10^{t_u}$, $t_u\simeq 17.64$, and intersect the lines at {\em black points}.}\label{fg3} \end{figure} In fact, all superdense configurations with $\rho_*>\rho^{OV}$ are metastable due to the acoustic vibrations \cite{dk63, ha65} propagating in stars with characteristic time $T_a=(\gamma\bar{\rho})^{-1/2}$ where $\gamma$ denotes a gravitational constant and $\bar{\rho}=M_{NS}/V_{NS}$ denotes an average density of NS of the total mass $M_{NS}$ and volume $V_{NS}$. E.g., if $\rho_*\simeq 10^{19}$ then according to \cite{ha65} $\bar{\rho}\simeq 0.25\cdot 10^{15}$ and finally we have $T_a\simeq 2\cdot 10^{-4}$s. Simple calculation by formula (\ref{k22}) with $T_2=T_a$ and ${\mathbb P}(A_*)=1$, gives a value of $A_*$ that provides to have at least one strangelet in core during the time $T_a$, i.e., $A_*\simeq 200$. Suggestions of superdense stars with core density above $\rho ^{OV}$ continue to appear in the literature \cite{gl97, pr03}. | 14 | 4 | 1404.6685 |
1404 | 1404.5773_arXiv.txt | { The ability of metal free gas to cool by molecular hydrogen in primordial halos is strongly associated with the strength of ultraviolet (UV) flux produced by the stellar populations in the first galaxies. Depending on the stellar spectrum, these UV photons can either dissociate $\rm H_{2}$ molecules directly or indirectly by photo-detachment of $\rm H^{-}$ as the latter provides the main pathway for $\rm H_{2}$ formation in the early universe. In this study, we aim to determine the critical strength of the UV flux above which the formation of molecular hydrogen remains suppressed for a sample of five distinct halos at $z>10$ by employing a higher order chemical solver and a Jeans resolution of 32 cells. We presume that such flux is emitted by PopII stars implying atmospheric temperatures of $\rm 10^{4}$~K. We performed three-dimensional cosmological simulations and varied the strength of the UV flux below the Lyman limit in units of $\rm J_{21}$. Our findings show that the value of $\rm J_{21}^{crit}$ varies from halo to halo and is sensitive to the local thermal conditions of the gas. For the simulated halos it varies from 400-700 with the exception of one halo where $\rm J_{21}^{crit} \geq 1500$. This has important implications for the formation of direct collapse black holes and their estimated population at z > 6. It reduces the number density of direct collapse black holes by almost three orders of magnitude compared to the previous estimates. } | Observations of quasars at z $>6$ reveal that supermassive black holes (SMBHs) of a few billion solar masses were assembled within the first billion years after the Big Bang \citep{2003AJ....125.1649F,2006AJ....131.1203F,2010AJ....139..906W,2011Natur.474..616M,2013ApJ...779...24V}. Their formation mechanisms in the juvenile Universe remain unknown. The potential progenitors of SMBHs include the remnants of Pop III stars \citep{2001ApJ...552..459H,2004ApJ...613...36H,2009ApJ...696.1798T,2012ApJ...756L..19W,2014ApJ...781...60H,2014ApJ...784L..38M}, dense stellar cluster \citep{2004Natur.428..724P,2008ApJ...686..801O, 2009ApJ...694..302D} and direct collapse of a protogalactic gas cloud into so-called direct collapse black holes (DCBHs) \citep{2002ApJ...569..558O,2003ApJ...596...34B,2006ApJ...652..902S,2006MNRAS.370..289B,2006MNRAS.371.1813L,2008MNRAS.391.1961D,2008arXiv0803.2862D,2010MNRAS.402.1249S,2010MNRAS.tmp.1427J,2010A&ARv..18..279V,2010ApJ...712L..69S,2011MNRAS.411.1659L,2012RPPh...75l4901V,2012arXiv1203.6075H,2013MNRAS.436.2301P,2013MNRAS.433.1607L,2013MNRAS.tmp.2526L,2013ApJ...774...64W,2013ApJ...771...50A,2013MNRAS.433.1556Y,2013MNRAS.436.2989L,2013A&A...560A..34W,2014MNRAS.tmp..564L,2014MNRAS.440.1263Y,2014MNRAS.440.2969L,2014arXiv1404.4630I,2014arXiv1403.1293V}. Pristine massive primordial halos of $\rm 10^{7}-10^{8}~M_{\odot}$ which formed in the early universe at z$=15-20$ are the potential cradles for these DCBHs. It is imperative that their halos remain metal free and cooling is mainly regulated by atomic line radiation instead of $\rm H_{2}$. These conditions may lead to a monolithic isothermal collapse where fragmentation is suppressed and a supermassive star of $\rm 10^{4}-10^{6}~M_{\odot}$ forms which later collapses into a black hole (i.e., DCBH). This scenario is supported by numerical simulations which show that fragmentation remains inhibited and massive objects may form \citep{2003ApJ...596...34B,2008ApJ...682..745W,2009MNRAS.393..858R,2011MNRAS.411.1659L,2013MNRAS.432..668L,2013MNRAS.430..588L,2013MNRAS.433.1607L}. Recently, the feasibility of this scenario has been explored via high resolution numerical experiments and it is found that $\rm \sim10^{5}~M_{\odot}$ objects can form \citep{2013MNRAS.436.2989L,2014MNRAS.439.1160R}. These results are consistent with theoretical predictions \citep{2008MNRAS.387.1649B,2010MNRAS.402..673B,2011MNRAS.414.2751B,2012ApJ...756...93H,2012MNRAS.421.2713B,2013A&A...558A..59S,2013ApJ...778..178H,2013ApJ...768..195W}. Depending on the mass accretion rates these studies suggest the formation of supermassive stars or quasi-stars (stars with BH at their center) as potential embryos of DCBHs \citep{2013A&A...558A..59S}. In primordial gas, trace amount of $\rm H_{2}$ can be formed via gas phase reactions in the early universe which then leads to cooling and star formation. The main channel for the formation of $\rm H_{2}$ is: \begin{equation} \mathrm{H + e^{-} \rightarrow H^{-} +} \gamma \end{equation} \begin{equation} \mathrm{ H + H^{-} \rightarrow H_{2} + e^-.}\\ \label{h21} \end{equation} Once the first generation of stars, so-called PopIII stars, are formed they produce UV flux, pollute the intergalactic medium with metals via supernova explosions, and lead to a second generation of stars known as PopII stars. The UV flux produced by these stellar populations either photo-dissociates the molecular hydrogen directly or photo-detaches electrons from H$^-$ which provides the main route for the formation of H$_{2}$ in primordial gas chemistry. The stellar spectra of PopIII stars are harder with a characteristic temperature of $\rm 10^{5}~K$ while PopII stars are characterized by soft spectra with temperatures of $\rm 10^{4}~K$. PopIII stars mainly contribute to the direct dissociation of H$_2$ while PopII stars also photo-detach H$^-$. UV photons with energies between 11.2 and 13.6 eV are absorbed in the Lyman-Werner bands of molecular hydrogen and put it into an excited state. The $\rm H_{2}$ molecule later decays to the ground state and gets dissociated as \begin{equation} \mathrm{ H_{2}} + \gamma \mathrm{ \rightarrow H_{2}^{*} \rightarrow H + H } , \\ \end{equation} a process known as the Solomon process. On the other hand, H$^-$ photo-detachment occurs via low energy photons above 0.76 eV. In this study, we focus on the background UV flux predominantly emitted by PopII stars as tiny amounts of metals can lead to fragmentation \citep{2005ApJ...626..627O,2013ApJ...766..103D}. The critical value of the UV flux, hereafter called $\rm J_{21}^{crit}$, above which $\rm H_{2}$ cooling remains suppressed, can be determined by comparing the $\rm H_{2}$ formation and dissociation time scales. \cite{2001ApJ...546..635O} found from one-zone calculations that $\rm J_{21}^{crit}=10^3$ in units of $\rm J_{21}=10^{-21}~erg~cm^{-2}~s^{-1}~Hz^{-1}~sr^{-1}$ for $\rm T_{*}=10^4~K$ which was later confirmed by \cite{Bromm03} in 3D simulations for a single halo. These estimates were revised by \cite{2010MNRAS.402.1249S} (hereafter S10) through three dimensional simulations using the $\rm H_{2}$ self-shielding formula of \cite{1996ApJ...468..269D}, finding that $\rm J_{21}^{crit}= 30-300$. They attributed these differences to the choice of a more accurate and higher $\rm H_{2}$ collisional dissociation rate, and focused on rather massive halos forming at z $<$10. \cite{2011MNRAS.418..838W} (hereafter WG11) improved the H$_2$ self-shielding function of \cite{1996ApJ...468..269D} and anticipated that it may further reduce the value of $\rm J_{21}^{crit}$. Such values of $\rm J_{21}^{crit}$ are much larger than the global background flux but can be achieved in the close vicinity (about 10 kpc) of nearby star forming galaxies \citep{2008MNRAS.391.1961D,2012MNRAS.425.2854A,2014arXiv1403.5267A}. In this article, we derive the values of $\rm J_{21}^{crit}$ for a stellar spectrum of $\rm T_{*}=10^4~K$ employing the improved $\rm H_{2}$ self-shielding fitting function provided by \cite{2011MNRAS.418..838W}. Major improvements compared to the previous studies are the following: \begin{itemize} \item Selection of a larger sample of halos with collapse redshifts at $z>$10. \item Employed higher order chemical solver DLSODES \citep{2013MNRAS.434L..36B}. \item Accurate determination of $\rm J_{21}^{crit}$ for the individual halos. \item Higher Jeans resolution of 32 cells. \item Improved self-shielding function of WG11. \end{itemize} We note that the importance of an accurate chemical solver in high resolution simulations was previously reported by \cite{2013MNRAS.434L..36B}. The impact of higher Jeans resolution has also been shown by \cite{2013MNRAS.430..588L} and \cite{2012ApJ...745..154T}. Our selected halos are collapsed at $z>$10 in contrast to S10 where halos collapsed at $z<$10. All these improvements distinguish the present work from S10. We perform three dimensional cosmological simulations for five different halos of a few times $\rm 10^{7}~M_{\odot}$ and vary the strength of the background UV flux (hereafter $\rm J_{21}$, i.e. UV flux below Lyman limit). We use the chemical network listed in table A1 of S10 which includes all the relevant process for the formation and dissociation of molecular hydrogen. We further employed a fixed Jeans resolution of 32 cells throughout the simulations for better resolving the shocks and turbulence. A particular goal of this paper is to provide a rather narrow constraint on $\rm J_{21}^{crit}$ for individual halos, and to point at potential correlations with halo properties. This study has important implications for the formation of DCBHs as it provides stronger estimates for the value of $\rm J_{21}^{crit}$ required for dissociation of molecular hydrogen. The organization of this article is as follows. In section 2, we provide the details of simulations setup and summary of a chemical network. In the third section, we present our findings and discuss our conclusions in section 4. | One of the main obstacles for the formation of direct collapse black holes is to avoid fragmentation in massive primordial halos which are the potential birthplaces of seed black holes. This may only be possible in the absence of molecular hydrogen which may induce fragmentation and trigger star formation. The ubiquity of background UV flux can photo-dissociate $\rm H_{2}$ molecules and may overcome this obstruction. The prime objective of this work is to determine the critical value of the background UV flux required to suppress the molecular hydrogen formation in atomic cooling halos. As photo-dissociation of $\rm H_{2}$ depends on the type of stellar spectrum, we have considered here UV photons below 13.6 eV emitted by PopII stars. We have conducted three-dimensional cosmological simulations for five distinct halos by including all relevant processes for the formation and dissociation of $\rm H_{2}$. The halos studied here have typical masses of a few times 10$^7$~M$_{\odot}$ and were illuminated by various strengths of background UV flux. We here employed the $\rm H_{2}$ self-shielding fitting function provided by WG11. Our findings show that the value of $\rm J_{21}^{crit}$ strongly depends on the properties of the halo and may vary from halo to halo. For the halos studied here, we found that the value of $\rm J_{21}^{crit}$ varies from 400-700 with the exception of one halo where it is about 1500. It is also found that $\rm J_{21}^{crit}$ may depend on the mass of halo, as the two most massive halos have reduced values of $\rm J_{21}^{crit}$. This trend is consistent with results by S10, where more massive halos at have even lower values of $\rm J_{21}^{crit}$. We note that our one-zone calculations are in agreement with S10. The highly non-linear collapse dynamics leads to the occurrence of shocks with Mach numbers of about 3 at various densities which changes the local gas temperature and consequently $\rm J_{21}^{crit}$ differs from halo to halo due to the strong dependence of H$_{2}$ collisional rate on temperature and density \citep{1996ApJ...461..265M}. To build up the supermassive black holes at z $~$6-7, one should preferentially consider halos that collapse early, implying a lower mass and potentially higher $\rm J_{21}^{crit}$ at the same virial temperature. However, it will be desirable in the future to verify it for a larger sample of halos with broader mass range. Our estimates for $\rm J_{21}^{crit}$ are quite robust as we consider a larger sample of halos and higher Jeans resolution leading to better resolved shocks and employed the high order chemical solver DLSODES. We also find that the value of $\rm J_{21}^{crit}$ weakly depends on the choice of $\rm H_{2}$ self-shielding. Although, the fitting formula of \cite{1996ApJ...468..269D} overestimates $\rm H_{2}$ self-shielding compared to WG11, its impact is very low for the adapted stellar spectra. The value of $\rm J_{21}^{crit}$ is about an order of magnitude higher in 3D calculations compared to the one-zone results. This is because of the inability of one-zone calculations to model shocks and hydrodynamical effects. Similar results have been found in the study of S10. We also included the effect of dissociative tunneling and found from the one-zone test that it decreases $\rm J_{21}^{crit}$ by a factor of three. The estimates of $\rm J_{21}^{crit}$ determined in this work have important implications for the formation of DCBHs. Our results suggest that the value of $\rm J_{21}^{crit}>400$ should be employed in computing the number density of DCBHs. The value of $\rm J_{21}^{crit}$ used in previous studies \citep{2012MNRAS.425.2854A,2014arXiv1403.5267A} seems rather low (i.e. 30) and may be one of the reasons for the high abundance of DCBHs predicted from semi-analytical calculations. From our results, the expected BH number density is ~$\rm 10^5$ per comoving $\rm Gpc^{-3}$ at z= 6 for $\rm J_{21}^{crit}=400$. This estimate is obtained by rescaling the values of \citep{2012MNRAS.425.2854A} for a higher $\rm J_{21}^{crit}$. These estimates do not take into account the metal enrichment in the intergalactic medium by supernova driven winds which may further reduce the number density of DCBHs. In our previous studies \citep{2013MNRAS.433.1607L,2013MNRAS.436.2989L,2014MNRAS.440.2969L}, we have shown that the presence of strong UV flux leads to an isothermal collapse where conditions are fertile for the formation of DCBHs. In fact, under these conditions large accretion rates of $>1$~M$_{\odot}$/yr are observed which result in the formation of supermassive stars of $\rm 10^{5}~M_{\odot}$, the potential progenitors of DCBHs. We have presumed here that these halos exposed to the intense UV flux by PopII stars are metal-free and remain pristine throughout their evolution. The transition from PopIII to PopII stars may inject metals and pollute these halos. Depending on the critical value of metallicity which can be as low as $\rm10^{-5}$ Z/Z$_{\odot}$ \citep{2009A&A...496..365C,2011ApJ...737...63A,2012A&A...540A.101L}, once the metal content in halos exceeds this value fragmentation becomes inevitable \citep{2008ApJ...686..801O}. Particularly, the cooling due to the dust even in the presence of a strong UV flux becomes effective at densities around $\rm 10^{12}-10^{15}~cm^{-3}$ and may lead to the formation of dense stellar clusters \citep{2009ApJ...694..302D}. Nevertheless, metal enrichment in the universe is expected to be patchy and pristine halos may exist down to $\rm z>6$. Given the strong dependence of $\rm J_{21}^{crit}$ on the local thermal conditions as argued above, the local heating/cooling effects such as heating by ambiploar diffusion, turbulence dissipation as well as cooling by HD molecules may change the critical value of the flux by a factor of few. In fact the impact of turbulence and magnetic field in the presence of UV flux was explored by \cite{2013A&A...553L...9V} and they found that in turbulent halos with stronger initial seed fields the value of $\rm J_{21}^{crit}$ is reduced by an order of magnitude. Furthermore, the presence of cosmic rays/X-rays may significantly enhance the critical value of $\rm J_{21}$ \citep{2011MNRAS.416.2748I}. It was recently pointed out by \cite{2014arXiv1403.6155R} that including the effect of turbulence in Doppler broadening reduces the $\rm H_{2}$ self-shielding. This should be explored in future studies. | 14 | 4 | 1404.5773 |
1404 | 1404.0360_arXiv.txt | Motivated by the idea that inflation occurs at the GUT symmetry breaking scale, in this paper we construct a new class of large field inflaton potentials where the inflaton starts with a power law potential; after initial period of relative fast roll that lasts until after a few {\it e-folds} inside the horizon, it transits to the attractor of the slow roll part of the potential with a lower power. Due to the initial fast roll stages of inflation, we find a suppression in scalar primordial power at large scales and at the same time the choice of the potential can provide us a tensor primordial spectrum with high amplitude. This suppression in scalar power with a large tensor-to-scalar ratio helps us to reconcile the Planck and BICEP2 data in a single framework. We find that a transition from a cubic to quadratic form of inflaton potential generates an appropriate suppression in power of scalar primordial spectrum that provides significant improvement in fit compared to power law model when compared with Planck and BICEP2 data together. We calculate the extent of non-Gaussianity, specifically, the bispectrum for the best fit potential and show that it is consistent with Planck bispectrum constraints. | 14 | 4 | 1404.0360 |
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1404 | 1404.0683_arXiv.txt | A polytropic model is used to investigate the effects of dark photospheric spots on the evolution and radii of magnetically active, low-mass ($M<0.5\,M_{\odot}$), pre-main sequence (PMS) stars. Spots slow the contraction along Hayashi tracks and inflate the radii of PMS stars by a factor of $(1-\beta)^{-N}$ compared to unspotted stars of the same luminosity, where $\beta$ is the equivalent covering fraction of dark starspots and $N \simeq 0.45 \pm 0.05$. This is a much stronger inflation than predicted by the models of Spruit \& Weiss (1986) for main sequence stars with the same $\beta$, where $N \sim 0.2$--0.3. These models have been compared to radii determined for very magnetically active K- and M-dwarfs in the young Pleiades and NGC~2516 clusters, and the radii of tidally-locked, low-mass eclipsing binary components. The binary components and ZAMS K-dwarfs have radii inflated by $\sim 10$ per cent compared to an empirical radius-luminosity relation that is defined by magnetically inactive field dwarfs with interferometrically measured radii; low-mass M-type PMS stars, that are still on their Hayashi tracks, are inflated by up to $\sim 40$ per cent. If this were attributable to starspots alone, we estimate that an effective spot coverage of $0.35 < \beta < 0.51$ is required. Alternatively, global inhibition of convective flux transport by dynamo-generated fields may play a role. However, we find greater consistency with the starspot models when comparing the loci of active young stars and inactive field stars in colour-magnitude diagrams, particularly for the highly inflated PMS stars, where the large, uniform temperature reduction required in globally inhibited convection models would cause the stars to be much redder than observed. | There is increasing evidence that the radii of fast rotating, magnetically active K- and M-dwarf stars are inflated relative to the predictions of evolutionary models. Measurements of eclipsing binaries suggest that radii can be 10--15 per cent larger than expected, at a given mass, for binary components with $M<0.7\,M_{\odot}$ (Lopez-Morales 2007; Morales et al. 2009; Torres 2013). The components of these close binary pairs are expected to be tidally locked and fast-rotating, hosting strong, dynamo-generated magnetic fields. It has been suggested that the relative increase in radius is associated with this magnetic activity (Lopez-Morales 2007), although it is difficult to find slowly rotating, magnetically inactive eclipsing binaries with which to test this hypothesis. The same comparison between observation and theory cannot easily be made for single stars since a direct measurement of their masses is not possible. However, it is possible to compare the radii and luminosities of nearby field stars with the predictions of evolutionary models. Boyajian et al. (2012b, hereafter BM12) reported the interferometric angular diameters of K- and M-dwarfs with precise Hipparcos parallaxes and used these to determine an empirical radius-luminosity relation for main sequence (MS) stars over the temperature range 3200--5500\,K. The radii of these relatively inactive field stars show satisfactory agreement with the predicted radii of evolutionary models for K- and early M-dwarfs (e.g. the BCAH98 model of Baraffe et al. 1998) but for later M-dwarfs (spectral types M2 to M4), models generally underestimate radii (by $\sim 8$ per cent in the case of the BCAH98 model; $\sim 5$ per cent for the Dartmouth models of Dotter et al. 2008). No such comparisons have been made for magnetically active stars since there are no very active field M-dwarfs close enough to allow a precise interferometric radius measurement. Jackson, Jeffries \& Maxted (2009) estimated the mean radii for groups of rapidly rotating, highly magnetically active, late K- and M-dwarfs in the young open cluster NGC 2516 (age $\simeq 140$\,Myr -- Meynet, Mermilliod and Maeder 1993) using the product of their rotation periods and projected equatorial velocities. They found that their mean radii at a given luminosity are larger than model predictions and also larger than the interferometric radii of magnetically inactive dwarfs. The discrepancy rises from $\sim 10$ per cent for late K-dwarfs to as much as $\sim 50$ per cent for M4 dwarfs that are nearly or fully convective. This result supports the suggestion that the radius inflation (at a given luminosity) is due to rotationally induced magnetic activity. Further measurements are required to correlate radius inflation with rotation, magnetic field strength and other magnetic activity indicators, but there is sufficient evidence of a causal link to have stimulated theoretical studies. Two mechanisms related to dynamo-generated magnetic fields have been suggested that could inflate the radii of K- and M-dwarfs: inhibition or stabilisation of convection (e.g. Mullan \& MacDonald 2001; Chabrier et al. 2007; Feiden \& Chaboyer 2012, 2013a,b; MacDonald \& Mullan 2013); or the effect of cool, magnetic starspots (Spruit 1982; Spruit \& Weiss 1986; Chabrier et al. 2007). Mullan and MacDonald (2001) modelled the effect of interior magnetic fields by modifying the Schwarzschild criterion and suppressing convection. Chabrier et al. (2007) inhibited the convective efficiency by artificially reducing the convective mixing-length parameter. Feiden \& Chaboyer (2012, 2013b) described a modification to the Dartmouth evolutionary code that takes account of the effect of magnetic field on the equation of state and on mixing-length theory. In their papers they give examples that produce the modest (10 per cent) observed radius inflation of the components of several eclipsing binaries of mass $0.4<M/M_{\odot}<1.0$, although the predicted surface magnetic fields are a few times higher than observational estimates. This explanation of radius inflation becomes less plausible for lower mass stars where convective heat transfer is more and more efficient. Either much larger changes in effective mixing length are required to produce even a 10 per cent radius inflation in nearly or fully convective stars (e.g. the binary CM Dra), or interior magnetic field strengths would need to approach 50\,MG to sufficiently stabilise the star against convection, which are considered too high to be physically plausible (Feiden \& Chaboyer 2013a). An alternative mechanism is the effect of starspots in reducing heat flux out of the photosphere. Spruit \& Weiss (1986, hereafter SW86), modelled the effect of starspots on zero age main sequence (ZAMS) stars in the mass range 0.2--1.9\,$M_{\odot}$. They argued that the radius, luminosity and temperature of the unspotted surface would adapt on the Kelvin-Helmholtz timescale of the convective envelope to achieve thermal equilibrium. For low-mass stars with deep convective envelopes they used a polytropic model to show that the effect of starspots is to reduce the luminosity of the star at near-constant radius and temperature of the unspotted photosphere, hence lowering the {\it effective} temperature. For higher mass stars, where the mass of the convective envelope is a small fraction of the stellar mass, SW86 used a numerical model to show that the effect of starspots would be to increase the temperature of the unspotted photosphere while the radius and luminosity are nearly unchanged. Which of these mechanisms is dominant is undecided and likely to depend on the fraction of the photosphere covered by dark starspots. The presence of starspots on magnetically active stars is a well documented phenomenon. Evidence for dark starspots and their associated magnetic fields came initially from observations of rotational modulation of broadband fluxes (e.g. Hall 1972, Eaton \& Hall 1979) but there is now a large literature that detects and investigates starspots using direct and indirect observational techniques such as Doppler imaging (Collier-Cameron \& Unruh 1994; Strassmeier 2002), spectroscopy of Zeeman-broadened lines (e.g. Marcy 1982, Johns-Krull \& Valenti 1996), and tomography using circularly polarized light (Zeeman Doppler Imaging - Semel 1989, Donati et al. 1997). However, the relevant question here is whether spots cover enough area of the surface to significantly reduce the stellar luminosity? Spot coverage and spot temperatures have been determined for very active G- and K-stars from their optical TiO absorption bands, indicating filling factors of 20--50 per cent, with temperature ratios between spotted and unspotted photosphere of 0.65--0.76 (O'Neal, Neff \& Saar 1998; O'Neal et al. 2004; O'Neal 2006). If starspots are present at similar filling factors on active K- and M-dwarfs then this could significantly reduce their luminosity at a given radius. MacDonald and Mullan (2013) compared measurements of (projected) radii measured on young active M-dwarfs in NGC\,2516 (Jackson et al. 2009) with their numerical models that include a prescription for magnetic suppression of convection and cool starspots. They found that, for models in which only the effects of spots are included, the mean radii of the coolest stars in the sample would require a spot coverage of more than 79 per cent of the photosphere. As they considered this implausible they rejected spots as the main cause of radius inflation. Their estimate of the increase in radii was made by comparison to a model for the radii of inactive stars and, as we show in this paper, a significantly lower estimate of filling factor is obtained if the radii of active stars are instead compared to an empirical radius-luminosity relation for inactive main sequence stars (e.g. from BM12). SW86 estimated the effect of starspots only for ZAMS stars, where the luminosity depends on the temperature of their nuclear burning cores. The cool M-dwarfs in NGC\,2516, which show the largest apparent increase in radius relative to inactive stars, are still in their pre-main sequence (PMS) phase. They are likely to be fully convective with luminosity produced chiefly by the release of gravitational energy, and might respond differently to starspots. In this paper we: (a) extend the work of SW86 to include the effect of starspots on low-mass PMS stars; (b) use the recently reported radius-luminosity relation of BM12 for inactive MS stars as a baseline against which to assess the increase in radius of active K- and M- dwarfs; and (c) compare these models with the published radii of eclipsing binaries, and with the radii of highly active K- and M-dwarfs, including a new analysis of the projected radii of stars in the young Pleiades cluster. We find that PMS stars are more inflated for a given level of spot coverage than ZAMS stars of a similar luminosity and that the spot coverage required to produce the measured radii is large but perhaps not excessive. Finally it is shown that, in contrast to models that suppress convective flux and uniformly lower the photospheric temperature, the spot model is able to simultaneously match the loci of very active stars in colour-magnitude diagrams. \begin{figure*} \centering \begin{minipage}[t]{0.95\textwidth} \centering \includegraphics[width = 150mm]{fig1.eps} \end{minipage} \caption{ Evolutionary tracks for, 0.2, 0.3, 0.4 and 0.5 solar mass stars over the age range 5 to 150Myr. The solid lines in the left hand plot show evolutionary tracks from the BCAH98 model of stellar evolution ($[$M/H$]$=0, Y=0.275, $L_{mix}$ =$H_P$). The filled circles show the evolutionary tracks predicted using a polytropic model of an unspotted star on the Hayashi track with constants adjusted to match approximately the BCAH98 tracks. The crosses show the effect of 30 percent coverage of dark starspots for the same polytropic model. The right hand plot shows the variation of radius with time. The lines show the BCAH98 model and the filled circles show results of the polyropic model of an unspotted star.} \label{fig1} \end{figure*} | Section 2 described a simple polytropic model of a PMS star with dark starspots covering a fraction $\beta$ of the stellar surface. The model is applicable to stars on the Hayashi track if: \begin{enumerate} \item The stars are fully convective, without a radiative or a degenerate core and their luminosity results principally from the rate of change of gravitational potential energy as the star contracts. \item The effect of starspots is to reduce the mean Rosseland opacity averaged over the surface of the star as, $\overline{\kappa} \propto (1-\beta )$, where $\beta$ is the equivalent filling factor of dark starspots that would produce the same reduction in total luminosity as the actual coverage of starspots at the actual starspot temperatures. \item The simple relationships governing the radius inflation and change in temperature at a given luminosity (Eqn.~10) assume that the starspots were formed at a much earlier time. This condition seems likely to be met in clusters like NGC~2516 and the Pleiades, because evidence for starspots (e.g. modulation of light curves) is routinely reported in much younger star forming regions (Herbst \& Mundt 2005; Irwin \& Bouvier 2009). \end{enumerate} The model is therefore restricted to a range of masses and ages which must be determined by reference to an evolutionary model for unspotted stars. In this paper the polytropic model is restricted to stars of mass $0.2 < M/M_{\odot} < 0.5$ for ages $\leq 120$\,Myr. In sections 3 and 4 the predicted increase in radius at a given luminosity as a function of spot coverage is compared with the measured radii of active K- and early M-dwarfs in NGC~2516 and the Pleiades and with the radii of short-period eclipsing binaries with known distances. In our work we have used the polytropic model to scale an empirical relationship between radius and luminosity determined from inactive (and presumably unspotted) field stars with small corrections for differences in age and metallicity. The use of an empirical relation is important; at low luminosities the empirical radius-luminosity relation shows radii 8 per cent higher than the BCAH98 model isochrones (see Fig.~5). If the radii of active stars were compared directly with these model radii it would lead to the inference of an erroneously high level of radius inflation for the lower mass stars. As an example, Macdonald \& Mullan (2013) rejected starspots as the sole cause of radius inflation in the same sample of NGC~2516 stars, on the basis that the lowest luminosity objects would require an excessive ($\beta$ =0.79) level of spot coverage {\it relative to their model} of magnetically inactive stars. However, we showed in Fig.~5 that this model significantly underpredicts the interferometrically measured radii of inactive stars by about 15 per cent. Our lower estimate of the required spot coverage ($\beta = 0.51\pm 0.04$) arises principally because of our comparison with an empirical radius-luminosity relation. Using the empirical radius-luminosity relation does not avoid all systematic uncertainties. There is a radius uncertainty of 2--3 per cent in the mean baseline empirical relationship that can only be improved by more and better measurements of the interferometric radii of inactive low-mass stars. However, this level of systematic uncertainty is small compared to the radius inflation inferred from the projected radii of very active low-mass stars (see Fig.~9 and Table~4). The level of spot coverage our model requires to solely account for the inflated radii of active stars is $\beta=0.35$--0.51. This is a little higher than the effective filling factors, $\beta$=0.13--0.41, determined for very active G and K-dwarfs from analysis of their TiO absorption bands by O'Neal (2006) but not significantly so. It is still largely unknown what spot coverage fractions and spot temperatures might exist on very active M-dwarfs or low-mass PMS stars, though a small value of $\beta$ seems unlikely given the very high filling factors of kilogauss magnetic field suggested by Zeeman measurements of M dwarfs with $\log N_{R} < -1$ (Reiners, Basri \& Browning 2009). Jackson \& Jeffries (2013) have shown that the generally small amplitude of $I$-band light curves of the low-mass NGC~2516 members is quite consistent with $\beta \simeq 0.5$ if the spots are small and scattered over a large portion of the stellar surface. The required spot coverage could be (much) lower if the inflated radii were also partly (or mostly) explained by the inhibition of convection due to a globally pervasive magnetic field (Macdonald \& Mullan 2013; Feiden \& Chaboyer 2012b, 2013). A relatively simple way to test the relative merits of the starspots versus globally inhibited convection scenarios is to compare the location of active stars and inactive field stars in various colour-magnitude diagrams (CMDs). Stauffer et al. (2003) pointed out that K-dwarfs in the Pleiades are nearly 0.5 mag sub-luminous compared to a MS isochrone in the M$_V$ vs (B-V)$_0$ CMD and also significantly bluer (or sub-luminous) compared to much less active members of the older Praesepe cluster. However, in the M$_V$ vs (V-K)$_0$ CMD the Pleiades members appear redder (or more luminous). How does this compare with the anticipated effects of magnetic inhibition of convection and/or starspots? Expansion driven by magnetic inhibition of convection would lead to a uniformly lower surface temperature (at a given luminosity) and should exclusively redden highly active stars relative to magnetically inactive stars in {\it both} CMDs. The effect of starspots depends on the spot temperature ratio. For darker starspots (with a temperature ratio $\le$ 0.8), the $V$-band flux density from the spotted area is a relatively small fraction (10 to 20 per cent) of the flux density from the unspotted photosphere. In this case the effect of starspots will be to shift magnetically active stars blueward in the $M_V$~vs $(B-V)_0$ CMD. The extent of reddening, if any, in the $M_V$~vs $(V-K)_0$ plane is less easy to predict since the spotted area would still make a significant contribution to the $K$-band flux. Figure 10 shows CMDs for our sample of fast rotating stars in NGC~2516 and the Pleiades compared to empirical colour-magnitude relations for magnetically inactive MS stars from BM12, corrected to the metallicity and mean age of the clusters. Results are transformed from the radius-luminosity relations of BM12 to the colour-magnitude plane using their $(B-V)_0$ and $(V-K)_0$ colour-temperature relations and the bolometric corrections in Fig.~8. Also shown in Fig.~10 are the predicted effects of: \begin{itemize} \item Magnetic inhibition of convection producing a uniform reduction in surface temperature. Curves are shown for a 10 percent increase in radius for MS stars and 30 per cent increase for PMS stars. \item A simplified, two temperature, model of a spotted star with 50 per cent coverage of starspots, assuming a uniform spot temperature ratio of 0.7 between the spotted and unspotted photosphere, giving $\beta$=0.40. Bolometric corrections for the spotted areas, which are outside the range of the colour-temperature relations of BM12, are taken from the BT-Settl model atmospheres (Allard et al. 2003). \end{itemize} Results in the $M_V$~vs $(B-V)_0$ plane support the interpretation of Stauffer et al. (2003), that stars in the Pleiades are blue shifted relative to older, less magnetically active, stars because of significant spot coverage, though the blueward shift of the Pleiades stars is less than produced by a spot model with $\beta=0.4$. However a model solely invoking global inhibition of convection to explain radius inflation appears to be ruled out by the observations. Results in the $M_V$~vs $(V-K)_0$ are less clear cut. The model of a spotted star (with spot temperature ratio of 0.7) predicts a small blueing of active {\it MS stars} relative to inactive MS stars, whereas active MS stars in the Pleiades actually show a slight reddening. However, for the PMS regime at lower luminosities, and with more radius inflation, the two models make drastically different predictions. The low-mass PMS stars in NGC~2516 lie close to the empirical magnetically inactive locus and this is in reasonabe agreement with the prediction of the $\beta=0.4$ spot model. However to explain a $\geq 30$ per cent radius inflation seen in these stars using global magnetic inhibition requires much lower surface temperatures at a given luminosity and results in a predicted locus that is {\it much} redder than observed. This either points to (unknown) problems in the indirect methods used for determing the radii of these low-mass PMS stars or indicates that global suppression of convection cannot be the sole cause of radius inflation and favours the starspot model. It should be stressed that the modelled effects of starspots on the CMDs are qualitative in nature being based on a simple two temperature description of a spotted star. The real situation is likely to be more complex with spots and associated plages producing a range of surface temperatures. | 14 | 4 | 1404.0683 |
1404 | 1404.5918_arXiv.txt | The existence of the cosmic ray Halo in our Galaxy has been discussed for more than half a century. If it is real it could help to explain some puzzling features of the cosmic ray flux: its small radial gradient, nearly perfect isotropy and the low level of the fine structure in the energy spectra of the various particles. All these features could be understood if: (a) the Halo has a big size (b) cosmic rays in the Halo have a unform spatial or radial distribution and (c) the cosmic ray density in the Halo is comparable or even higher than that in the Galactic Disk. The main topic of the paper concerns the present status of the anisotropy and a model for its formation. In our model the extremely small amplitude of the dipole anisotropy is due to the dilution of the anisotropy in the Disk by the dominating isotropic cosmic rays from the Halo. Some minor deviations from complete isotropy in the sub-PeV and PeV energy regions point out to the possible contribution of the Single Source with the phase of its first harmonic opposite to the phase produced by the Disk. | The observed cosmic rays (CR) have several puzzling features which need to be explained and these are now listed: (a) Firstly, there is only a small radial gradient in the Galactic Disk (GD) in contrast with expectation, the reason is as follows. The most viable theory of the CR origin is that they are generated in the supernova (SN) explosions and acclerated by the shock waves in the supernova remnants (SNR) \cite{Ginz1}. According to recent studies \cite{Case,Green} the Galactocentric radial distribution of SNR is such that they are mostly concentrated in the Inner Galaxy with the maximum GD surface density at a Galactocentric radius of about $R\approx 3-4$ kpc followed by a rapid decrease at larger $R$. According to the model calculations \cite{EW1} the radial CR gradient coincides with that of SNR and at the Sun, of radial distance of 8.3 kpc, it should be equal to $S = d(lnI)/dR = (-0.17\pm 0.05)$ kpc$^{-1}$. Here $I$ is the CR intensity or the SNR density. However, the experimental values for the Outer Galaxy derived by us from the gamma-ray emissivity profile are $(-0.05\pm 0.03)$ kpc$^{-1}$ for both the second and third quadrants \cite{Abdo,Acker}, so that the observed CR radial distribution is significantly flatter than that expected from the distribution of their proposed sources. (b) Secondly, there is the surprising near-isotropy of the CR arrival directions. In the sub-PeV region the CR intensity is relatively high and allows the collection of good statistics with detectors of a reasonable size during an acceptable time. Theoretical calculations predict a the slow rise for the amplitude {\em A} of the first harmonic with energy {\em E} as $A \sim E^{0.3-0.5}$. On the opposite side the experimental measurements indicate a decreasing amplitude above a few TeV with a minimum of about $A\sim 2\cdot 10^{-4}$ for $E\sim (0.1-0.3)$PeV \cite{Guill,EW2}. The attempt to explain this decrease by an accidental spatial configuration of the sources is difficult because it shows that the probability of such a favorable configuration is definitely lower than a few percent \cite{EW3}. (c) The third puzzle is connected with the observed shape of the CR energy spectrum. Due to the stochastic distribution of the SNR in space and time there should be fine structure in the spectrum at some level. All realistic simulations confirm the possible existence of such structures \cite{EW1}. However, so far, only two structures are firmly established: the so called 'knee' at 3-4 PeV and the 'ankle' at 3-4 EeV. Below the knee, and between the knee and the ankle, measurements show quite a regular power law shape of the spectrum with only minor structure. In the last decade due to improvements in the energy resolution and increased statistics several works find hints of fine structure both below \cite{ATICn,CREAM, PAMELAn} and above the knee \cite{KASCADE1,GAMMA,TUNKA,IceTop,Yakutsk1}. In the region below the knee the experiments indicate a possible flattening of the proton and nuclei spectra above a rigidity of 200 GV. A similar flattening was also found in the primary electron plus positron spectrum \cite{ATICe1,ATICe2,PAMELAe}. However, the latest precise data from AMS-02 experiment do not show such features, except for positrons, in the same energy region below the knee \cite{AMS2}. The clarification of the situation is the duty of the experimental groups, but in any case the discussed irregularities, if they exist, are relatively small and do not disprove the basic feature of the CR energy spectrum, viz its nearly perfect regular power law shape. The present paper is an attempt to find a reasonable explanation of this puzzle, with special attention given to the large-scale anisotropy. | We develop a model which helps to give an explanation for at least one of the three puzzles mentioned in the Introduction, viz. the small radial gradient of the CR intensity, the small magnitude and peculiar energy dependence of the CR anisotropy and the small level of irregularities in the CR energy spectrum. In this paper we analyse the second puzzle - the anisotropy. The model exploits three basic ingradients: Disk, Halo and the Single Source. We postulate the dominant role of the Halo and its CR in our observations in spite of the fact that we are located inside the Disk. At PeV energies, approaching the knee, contributions from the Single Source begins to play an important role and inspired by the experimental evidence we assume that the phase of the CR intensity from the Single Source is opposite to the phase of the background CR from the Disk and the Halo. Due to this effect the amplitude of the dipole anisotropy decreases and approaches a minimum (dip) at sub-PeV energies. After that the amplitude begins to rise again, but because CR are mainly from the Single Source they come preferentially from the direction opposite to that at lower sub-PeV energies. The position and the shape of the dip is extremely sensitive to the parameters of the spectra adopted for the three ingredients: Disk, Halo and the Single Source. This sensitivity can be used for the study of the CR origin in the vicinity of the knee in the PeV energy region. \\ {\bf Acknowledgements} The authors are grateful to the Kohn Foundation for financial support. The (unknown) reviewer is thanked for very helpful comments and suggestions. | 14 | 4 | 1404.5918 |
1404 | 1404.4633_arXiv.txt | {Using the SofI instrument on the 3.5~m New Technology Telescope, we have conducted an extensive near-infrared monitoring survey of an unbiased sample of 69 brown dwarfs spanning the L0 to T8 spectral range, with at least one example of each spectral type. Each target was observed for a 2 -- 4 hour period in the $J_{\rm s}$-band, and the median photometric precision of the data is $\sim$~0.7\%. A total of 14 brown dwarfs were identified as variables with min-to-max amplitudes ranging from 1.7\% to 10.8\% over the observed duration. All variables satisfy a statistical significance threshold with a $p$-value $\leq 5$\% based on comparison with a median reference star light curve. Approximately half of the variables show pure sinusoidal amplitude variations similar to 2MASSJ2139+0220, and the remainder show multi-component variability in their light curves similar to SIMPJ0136+0933. It has been suggested that the L/T transition should be a region of a higher degree of variability if patchy clouds are present, and this survey was designed to test the patchy cloud model with photometric monitoring of both the L/T transition and non-transition brown dwarfs. The measured frequency of variables is $13^{+10}_{-4}$\% across the L7 -- T4 spectral range, indistinguishable from the frequency of variables of the earlier spectral types ($30^{+11}_{-8}$\%), the later spectral types ($13^{+10}_{-4}$\%), or the combination of all non-transition region brown dwarfs ($22^{+7}_{-5}$\%). The variables are not concentrated in the transition, in a specific colour, or in binary systems. Of the brown dwarfs previously monitored for variability, only $\sim60$\% maintained the state of variability (variable or constant), with the remaining switching states. The 14 variables include nine newly identified variables that will provide important systems for follow-up multi-wavelength monitoring to further investigate brown dwarf atmosphere physics.} | The L, T, and Y-type brown dwarfs represent a link between the coolest stars and giant planets. Many brown dwarfs are even cooler than currently observable exoplanetary atmospheres (e.g. HR8799b, HD189733b; \citealt{barman11}, \citealt{sing09,sing11}). The recently discovered Y dwarfs \citep{cushing11} approach the temperature of Jupiter. Since brown dwarfs never achieve a stable nuclear burning phase, they cool throughout their lifetimes, and temperature, rather than mass, is the dominant factor in defining the spectral sequence. As they cool, their atmospheres undergo changes in the chemistry and physical processes that sculpt their emergent spectra. While spectroscopy can be used to investigate atmospheric constituents and chemistry, photometric monitoring is an effective means to search for evidence of surface brightness inhomogeneities caused by cloud features, storms, or activity.\blfootnote{Based on observations made with ESO Telescopes at La Silla Observatory under programme ID 188.C-0493.} The transition region from late-L to early-T encompasses a particularly interesting change in physical properties, as the atmospheres transform from dusty to clear over a narrow effective temperature range, and the observed infrared colours reverse from red to blue. This is predicted to be an effect of the formation and eventual dissipation of dusty clouds in brown dwarf atmospheres \citep{chabrier00,marley02,burrows06}. Broadly, as brown dwarfs cool through the spectral sequence, the lower temperatures allow more complex molecules to form, resulting in condensate clouds. When the temperature is cool enough, large condensate grains cannot remain suspended high in the atmosphere and sink below the observable photosphere, allowing methane and molecular hydrogen to become the dominant absorbers. Although there are several existing models for condensate cloud evolution, most cannot easily explain the rapid colour change from red to blue over the L-to-T transition. A systematic survey of variability in brown dwarfs including both L/T transition objects and comparison hotter/cooler objects is required to search for differences in the structure of condensate clouds in this important regime. Existing photometric monitoring campaigns of brown dwarfs have been conducted at different wavelengths: optical bands (e.g. \citealt{tinney99} and \citealt{koen13}), near-IR bands (e.g. \citealt{artigau03}, \citealt{khandrika13} and \citealt{buenzli13}), mid-IR \citep[e.g.][]{morales-calderon06}, and radio frequencies \citep[e.g.][]{berger06}. From small ($<20$ objects) initial samples of ultracool field dwarfs, frequencies of variables ranged from 0\% to 100\% \citep[e.g. summary in][]{bailer-jones05}, and results from larger studies ($\sim25$ objects) have measured the frequency of variables to be in the range of 20\% to 30\% \citep[e.g.][]{khandrika13, buenzli13}. Examples of objects that vary in multiple wavebands have been identified (e.g. 2MASS J22282889-4310262 \citealt{clarke08, buenzli12}, SIMP J013656.5+093347.3 \citealt{artigau09}, 2MASS J21392676+0220226 \citealt{radigan12}), as well as objects recorded as variable in one wavelength range, but not another (e.g. 2MASS J15344984-2952274, \citealt{koen04b}). A small set of variable sources have been monitored contemporaneously at multiple wavelengths, with the combined results being used to infer the vertical extent of atmospheric features and to investigate atmospheric circulation patterns \citep[e.g.][]{buenzli12}. Given the unique probe of the atmospheric structure that multi-wavelength observations provide, it is essential to identify a larger set of known variables across a broad range of effective temperatures. Most monitoring programs have involved observation sequences spanning a few hours, but some studies have searched for longer timescale variations \citep[e.g.][]{gelino02,enoch03}. A time scale of a few hours is well-matched to a search for rotation-modulated variability, since expected rotation periods are $\sim2$ -- 12 hours for L and T dwarfs, considering the range of measured $v\sin i$ values (10 -- 60~km/s for L dwarfs and 15 -- 40~km/s for T dwarfs -- \citealt{zapatero06}) and the $\sim$~0.08 -- 0.10~$M_{\odot}$ radius of these objects from evolutionary models at the age of the field \citep{baraffe03}. Periodogram analysis of some variables has shown clear peaks associated with periods in the range of $\sim2$ -- 8~hours \citep[e.g.][]{clarke08,radigan12} which is consistent with an atmospheric feature rotating into and out of view. Other variables exhibit multi-component light curves \citep[e.g.][]{artigau09} that are suggestive of a rapid evolution of atmospheric features. To investigate the variability of brown dwarfs across the full L-T spectral sequence, we have performed a large-scale $J_{\rm s}$-band photometric monitoring campaign of 69 field brown dwarfs with the SofI instrument on the 3.5\,m New Technology Telescope (NTT). This survey is a part of the BAM (Brown dwarf Atmosphere Monitoring) project. In Section \ref{sample}, the properties of the sample, including magnitudes, spectral types, and companions are summarised. Details of the observations are reported in Section \ref{obs}, followed by the data reduction procedure, and methodology used to characterise each target as variable or constant in Section \ref{reduction}. Section~\ref{results} presents the results of the program and a comparison to previous variability studies. Finally, we discuss the sensitivity of the BAM survey and investigate possible correlations between variability and various observables such as spectral type, colour and binarity in Section~\ref{discussion}. The results are summarized in Section~\ref{conclusion}. | \label{discussion} \subsection{The sensitivity of the BAM survey} To obtain an estimate of the variability frequency for brown dwarfs across spectral types, it is essential to quantify the sensitivity of the data to detecting different amplitudes of variability. We estimate the sensitivity to variables of a certain amplitude as three times the target photometric uncertainty of each final target light curve. This places a limit on the minimum amplitude required for a detection above a certain statistical significance threshold. The proportion of the sample that is sensitive to a given variability amplitude is shown as a function of amplitude in Figure~\ref{sensitivity}. As shown in Figure~\ref{sensitivity}, the BAM survey is capable of detecting any object in the sample showing a peak-to-trough amplitude $\ge2.3\%$ during the duration of the observations. The detection probability continues to decrease with decreasing amplitude with a sensitivity of 50\% occurring for variables with a $\sim1.7$\% amplitude. Given that the full BAM sample is sensitive to variables with amplitudes $\ge2.3\%$, Table~\ref{tabl:AmpLimits} quantifies the frequency of variability for different subsets of spectral types using an amplitude cutoff of 2.3\% and $p$-value $\le0.05$; this level includes all but one BAM $p\le0.05$ variable. Figure~\ref{var_fraction} shows how the variability frequency (considering all spectral types) varies as a function of amplitude to account for the declining proportion of the sample that is sensitive to lower amplitude variables. To calculate the uncertainty on the variability frequency we use the binomial distribution \begin{table*} \tiny{\centering \caption{Limits on constant targets in this survey.} \label{tabl:constants} \begin{tabular}{lcccccccc} \hline Object & Spectral Type & Obs. Dur. [hours] & Refs. & DOF & $\chi^2_\nu$ & $\tilde{\eta}$ & $Q$ (\%) & $p$-value (\%)\\ \hline \hline 2MASS J00165953-4056541 & L3.5 & 3.43 & 6 & 8 & 0.9 & 0.8 & 0.54 & 51.1\\ 2MASS J00184613-6356122 & L2 & 3.43 & 7 & 8 & 1.1 & 1.0 & 0.42 & 35.4\\ 2MASS J00345157+0523050 & T6.5 & 2.98 & 7 & 7 & 0.9 & 1.1 & 0.53 & 49.1\\ 2MASS J01282664-5545343 & L1 & 2.66 & 5 & 5 & 0.5 & 2.1 & 0.61 & 77.7\\ 2MASS J02284355-6325052 & L0 & 2.84 & 3 & 10 & 0.8 & 1.0 & 0.42 & 62.7\\ 2MASS J02572581-3105523 & L8 & 3.33 & 7 & 5 & 0.3 & 1.2 & 0.48 & 90.2\\ 2MASS J03185403-3421292 & L7 & 1.97 & 5 & 4 & 1.9 & 1.8 & 0.99 & 10.1\\ 2MASS J03400942-6724051 & L7 & 2.61 & 6 & 5 & 0.2 & 1.1 & 0.75 & 96.4\\ 2MASS J04070752+1546457 & L3.5 & 3.19 & 7 & 7 & 0.7 & 0.7 & 0.43 & 68.4\\ 2MASS J04151954-0935066 & T8 & 3.29 & 4 & 7 & 0.3 & 0.6 & 0.83 & 95.8\\ SDSS J042348.56-041403.5 & T0 & 2.62 & 8 & 5 & 0.4 & 0.7 & 0.51 & 86.2\\ 2MASS J04455387-3048204 & L2 & 4.55 & 7 & 9 & 0.7 & 0.7 & 0.36 & 68.7\\ 2MASS J05103520-4208140 & T5 & 2.74 & 6 & 6 & 0.5 & 0.6 & 0.66 & 84.2\\ 2MASS J05160945-0445499 & T5.5 & 3.11 & 4 & 7 & 1.1 & 1.1 & 0.88 & 38.5\\ 2MASS J05233822-1403022 & L5 & 4.56 & 6 & 9 & 0.7 & 1.0 & 0.99 & 66.8\\ 2MASS J05591914-1404488 & T4.5 & 3.11 & 6 & 7 & 0.3 & 0.5 & 0.37 & 96.3\\ 2MASS J06244595-4521548 & L5 & 3.26 & 5 & 6 & 1.2 & 1.0 & 0.49 & 28.5\\ 2MASS J07290002-3954043 & T8 & 3.37 & 8 & 10 & 0.6 & 0.6 & 0.9 & 84.3\\ DENIS J081730.0-615520 & T6 & 3.48 & 8 & 13 & 0.8 & 0.7 & 0.68 & 64.5\\ 2MASS J09153413+0422045 & L7 & 3.33 & 6 & 7 & 0.9 & 0.8 & 0.55 & 50.1\\ 2MASS J09393548-2448279 & T8 & 3.06 & 8 & 9 & 1.0 & 0.8 & 0.98 & 46.5\\ 2MASS J09490860-1545485 & T2 & 3.16 & 8 & 6 & 1.7 & 1.3 & 0.67 & 11.0\\ 2MASS J10043929-3335189 & L4 & 3.14 & 8 & 11 & 1.4 & 1.0 & 0.99 & 17.6\\ 2MASS J10073369-4555147 & T5 & 3.41 & 8 & 13 & 0.3 & 0.5 & 1.62 & 99.5\\ 2MASS J10210969-0304197 & T3 & 3.17 & 7 & 6 & 0.7 & 0.6 & 0.88 & 64.1\\ 2MASS J11145133-2618235 & T7.5 & 2.90 & 5 & 5 & 0.4 & 1.0 & 1.0 & 86.3\\ 2MASS J11555389+0559577 & L7.5 & 2.87 & 5 & 7 & 0.6 & 1.2 & 0.74 & 79.3\\ 2MASS J12255432-2739466 & T6 & 2.85 & 7 & 5 & 0.4 & 0.3 & 0.67 & 86.9\\ 2MASS J12281523-1547342 & L6 & 3.39 & 7 & 10 & 1.0 & 1.1 & 0.5 & 42.5\\ 2MASS J12314753+0847331 & T5.5 & 2.56 & 5 & 5 & 0.3 & 1.6 & 1.63 & 89.6\\ 2MASS J12545393-0122474 & T2 & 3.17 & 7 & 9 & 1.0 & 1.1 & 0.51 & 45.2\\ 2MASS J13262981-0038314 & L5.5 & 2.80 & 4 & 7 & 0.9 & 0.9 & 1.17 & 50.8\\ 2MASS J14044941-3159329 & T2.5 & 3.01 & 8 & 8.5 & 0.4 & 0.6 & 0.82 & 89.4\\ 2MASS J15074769-1627386 & L5.5 & 4.16 & 7 & 15 & 0.4 & 0.6 & 1.21 & 98.5\\ SDSS J151114.66+060742.9 & T.0 & 3.85 & 7 & 11 & 0.8 & 0.6 & 0.8 & 64.1\\ 2MASS J15210327+0131426 & T2 & 6.37 & 7 & 12 & 0.5 & 0.6 & 0.9 & 93.6\\ 2MASS J15344984-2952274 & T5.5 & 3.88 & 8 & 8 & 0.5 & 0.6 & 0.62 & 85.3\\ 2MASS J15462718-3325111 & T5.5 & 3.50 & 4 & 7 & 0.4 & 0.5 & 1.45 & 93.3\\ 2MASS J15530228+1532369 & T7 & 3.82 & 7 & 8 & 2.7 & 1.0 & 0.72 & 0.7\\ 2MASS J16241436+0029158 & T6 & 3.21 & 8 & 8 & 0.3 & 0.4 & 0.66 & 95.6\\ 2MASS J16322911+1904407 & L8 & 3.89 & 4 & 14 & 1.1 & 1.0 & 1.76 & 36.4\\ 2MASS J18283572-4849046 & T5.5 & 3.06 & 8 & 7 & 0.4 & 0.5 & 0.5 & 88.7\\ 2MASS J19360187-5502322 & L5 & 3.21 & 7 & 6 & 0.4 & 0.5 & 0.4 & 86.9\\ SDSS J204317.69-155103.4 & L9.0 & 3.03 & 8 & 6 & 1.4 & 1.0 & 1.14 & 22.5\\ SDSS J204749.61-071818.3 & T0.0 & 2.68 & 8 & 5 & 1.6 & 1.2 & 1.16 & 15.8\\ 2MASS J20523515-1609308 & T1 & 3.08 & 8 & 7 & 0.8 & 0.8 & 0.7 & 62.9\\ 2MASS J21513839-4853542 & T4 & 3.08 & 8 & 7 & 0.8 & 0.7 & 0.51 & 63.0\\ 2MASS J22521073-1730134 & L7.5 & 2.34 & 5 & 5 & 0.8 & 1.3 & 0.65 & 58.9\\ ULAS J232123.79+135454.9 & T7.5 & 2.94 & 8 & 7 & 0.3 & 0.6 & 0.9 & 94.9\\ 2MASS J23224684-3133231 & L0 & 2.69 & 6 & 4 & 1.1 & 0.9 & 0.49 & 37.8\\ 2MASS J23312378-4718274 & T5 & 2.63 & 5 & 6 & 1.7 & 1.1 & 1.13 & 12.0\\ 2MASS J23565477-1553111 & T6 & 2.98 & 5 & 7 & 3.5 & 0.6 & 0.64 & 0.1\\ \hline \end{tabular} } \end{table*} \begin{table*} \tiny{\centering \caption{Summary of variable sources.} \label{tabl:variable_sample} \begin{tabular}{lcccl} \hline \multicolumn{1}{c}{Target Name} & Band & Variable/Constant & References & \multicolumn{1}{c}{Notes} \\ \hline \hline \multicolumn{5}{c}{New variables from this study with no prior observations} \\ \hline 2MASS J00501994-3322402 & $J_{\rm s}$ & V & & \\ 2MASS J01062285-5933185 & $J_{\rm s}$ & V & & \\ 2MASS J03582255-4116060 & $J_{\rm s}$ & V & & \\ 2MASS J09310955+0327331 & $J_{\rm s}$ & V & & Candidate Variable ($p$-val$ = 6.1$\%) \\ 2MASS J10101480-0406499 & $J_{\rm s}$ & V & & \\ 2MASS J12074717+0244249 & $J_{\rm s}$ & V & & \\ 2MASS J22551861-5713056 & $J_{\rm s}$ & V & & \\ \hline \multicolumn{5}{c}{New variables previously categorised as constant} \\ \hline DENIS J0205.4-1159 & $I_{\rm c}$ & C & K13 & \\ Candidate Variable ($p$-val$ = 7.9$\%) & $K_{\rm s}$ & C & E03 & $<3$\%\\ & 8.46 GHz & C & B06 & $< 30 \mu\rm{Jy}$\\ 2MASS J03480772-6022270 & $J$ & C & C08 & $<10$~mmag, periodic \\ 2MASS J04390101-2353083 & $I_{\rm c}$ & C & K13 & \\ & $I_{\rm c}$ & C & K05 & \\ & 8.46 GHz & C & B06 & $<$42$\mu$Jy\\ 2MASS J11263991-5003550 & $I_{\rm c}$ & C & K13 & possibly periodic \\ \hline \multicolumn{5}{c}{Literature variables confirmed as variable in this study} \\ \hline SIMP J013656.5+093347.3 & $JK$ & V & A09, Ap13 & $\Delta J=4.5\%$, P = 2.39 $\pm$ 0.05 hr\\ 2MASS J08354256-0819237 & $I_{\rm c}$ & V & K04, K13 & $\Delta I_{\rm c}$=10-16~mmag, P=3.1~hr \\ & 8.46 GHz & C & B06 & $<30\mu\rm{Jy}$\\ 2MASS J12171110-0311131 & $J$ & V & A03 & $\Delta J = 0.176 \pm 0.013$~mag \\ Candidate Variable (p-val$ = 9.3$\%) & 8.46 GHz & C & B06 & $<111\mu\rm{Jy}$\\ & & & Z06 & \\ 2MASS J13004255+1912354 & $I_{\rm c}$ & C & K13 & \\ & $I$ & V & G02 & P=238$\pm$9hr\\ & 8.46 GHz & C & B06 & $<87\mu\rm{Jy}$\\ & $J+K$' & C & Kh13 & $J <$ 1.1\%, $K$' $<$ 1.7\%\\ & $I_{\rm c}$ & C & K05 & \\ & $JHK_{\rm s}$ & C & KMM04 &\\ 2MASS J21392676+0220226 & $JHK_{\rm s}$ & V & R12, Ap13 & $\Delta$($J$,$H$,$K$)=(0.3, 0.18, 0.17)~mag, P = 7.721$\pm$0.005~hr \\ & $J+K$' & V & Kh13 & $\Delta J=6.7\%$, C at $K$' $<8$\%\\ 2MASS J22282889-4310262 & $J$ & V & C08, Bu12 & $\Delta J=15.4 \pm 1.4$~mmag, P=1.43$\pm$0.16hr\\ & 8.46 GHz & C & B06 & $<30\mu\rm{Jy}$\\ \hline \multicolumn{5}{c}{Objects with reported IR variability measured as constants in this study} \\ \hline SDSS J042348.56-041403.5 & $I_{\rm c}$ & C & K13 & \\ & $K_{\rm s}$ & likely V & E03 & $0.3\pm0.18$~mag, P$=1.39$~--~$1.62$~hr\\ & $J$ & V & C08 & $8.0\pm0.8$~mmag, P$=2\pm0.4$hr\\ & $JHK_{\rm s}$ & C & KTTK05 & $J$ $<15$~mmag, $H< 11$~mmag, $K <2$~mmag\\ & 8.46 GHz & C & B06 & $<42\mu\rm{Jy}$\\ 2MASS J05591914-1404488 & HST G141 grism & V & Bu13 & \\ & $I_{\rm c}$ & C & K13 & \\ & $K_{\rm s}$ & C & E03 & $<7$\%\\ & $J$ & C08 & C & $<5$~mmag\\ & $I_{\rm c}$ & C & K04 & \\ & 8.46 GHz & B06 & $<27\mu\rm{Jy}$\\ 2MASS J06244595-4521548 & HST G141 grism & V & Bu13 & \\ DENIS J081730.0-615520 & HST G141 grism & V & Bu13 & \\ 2MASS J09393548-2448279 & $K$' & V & Kh13 & 0.31~mag\\ & $J$ & C & & $<0.141$~mag\\ 2MASS J15344984-2952274 & $I_{\rm c}$ & C & K13 & \\ & $I_{\rm c}$ & C & K05 & \\ & $JHK_{\rm s}$ & C & KTTK05 & $J <10$~mmag, $H <11$~mmag, $K <$18~mmag\\ & $JHK_{\rm s}$ & V & KMM04 & $H$ $4$~mmag, $K$ 7~mmag, P$=0.96$~hr\\ & 8.46 GHz & C & B06 & $<63\mu\rm{Jy}$\\ 2MASS J16241436+0029158 & HST G141 grism & V & Bu13 & Variability detected in water band (1.35-1.44~$\mu m$)\\ & $JHK_{\rm s}$ & C & KMM04 & \\ & 8.46 GHz & C & B06 & $<$36$\mu$Jy\\ 2MASS J23312378-4718274 & $J$ & V & C08 & $\Delta J=12.4 \pm 1.3$~mmag, P=2.9$\pm$0.9~hr \\ \hline \multicolumn{5}{c}{Objects with reported Optical variability measured as constants in this study} \\ \hline 2MASS J02284355-6325052 & $I_{\rm c}$ & V & K13 & \\ \hline \end{tabular} References: \citet{artigau03} [A03], \citet{artigau09} [A09], \citet{apai13} [Ap13], \citet{berger06} [B06], \citet{buenzli12} [Bu12], \citet{buenzli13} [Bu13], \citet{clarke08} [C08], \citet{enoch03} [E03], \citet{gelino02} [G02], \citet{khandrika13} [Kh13], \citet{koen04a} [K04], \citet{koen04b} [KMM04], \citet{koen05b} [KTTK05], \citet{koen05a} [K05], \citet{koen13} [K13], \citet{radigan12} [R12], \citet{zapatero06} [Z06]\\ } \end{table*} \begin{table} \tiny{ \centering \caption{Summary of constant sources.} \label{tabl:LitConst} \begin{tabular}{lccl} \hline \hline \multicolumn{1}{c}{Target Name} & Band & References & \multicolumn{1}{c}{Notes} \\ \hline 2MASS J02572581-3105523 & $I_{\rm c}$ & K13 & \\ 2MASS J04070752+1546457 & $JK'$ & Kh13 & no results\\ 2MASS J04151954-0935066 & 8.46 GHz & B06 & $<\mu\rm{Jy}$\\ 2MASS J04455387-3048204 & $I_{\rm c}$ & K13 & \\ & $I_{\rm c}$ & K04 & \\ & 8.46 GHz & B06 & $<66\mu\rm{Jy}$\\ 2MASS J05233822-1403022 & $I_{\rm c}$ & K13 & \\ & $I_{\rm c}$ & K05 & \\ & 8.46 GHz & B06 & $231\pm14\mu\rm{Jy}$\\ 2MASS J12255432-2739466 & $J$ & KTTK05 & $<12$~mmag\\ & $H$ & & $<14$~mmag\\ & $K_{\rm s}$ & & $<9$~mmag\\ & $JHK_{\rm s}$ & KMM04 & \\ 2MASS J12281523-1547342 & $I_{\rm c}$ & K13 & \\ & 8.46 GHz & B06 & $<87\mu\rm{Jy}$\\ 2MASS J12545393-0122474 & $JHK_{\rm s}$ & KMM04 & \\ & $J$ & Gi13 & $<5$mmag \\ 2MASS J15074769-1627386 & $I_{\rm c}$ & K13 & \\ & $I_{\rm c}$ & K03 & \\ & 8.46 GHz & B06 & $<57\mu\rm{Jy}$\\ 2MASS J1511145+060742 & $J$ & Kh13 & $<3.3$\%\\ & $K$'& & $<6.1$\%\\ 2MASS J15462718-3325111 & $JHK_{\rm s}$ & KMM04 & \\ 2MASS J15530228+1532369 & $JHK_{\rm s}$ & KMM04 & \\ 2MASS J16322911+1904407 & 8.46 GHz & B06 & $<54\mu$Jy\\ & HST G141 Grism & Bu13 & \\ 2MASS J19360187-5502322 & $I_{\rm c}$ & K13 & \\ 2MASS J23224684-3133231 & $I_{\rm c}$ & K13 & \\ \hline \end{tabular} References: \citet{berger06} [B06], \citet{buenzli13} [Bu13], \citet{girardin13} [Gi13], \citet{khandrika13} [Kh13], \citet{koen04a} [K04], \citet{koen03} [K03], \citet{koen04b} [KMM04], \citet{koen05b} [KTTK05], \citet{koen05a} [K05], \citet{koen13} [K13].\\ } \end{table} \begin{figure} \centering \includegraphics[width=84mm]{Figure8.pdf} \caption{Proportion of the survey sensitive to variability as a function of peak-to-trough amplitudes for different detection thresholds. The dashed line represents the fraction of objects with a photometric accuracy good enough to have allowed for the detection of variability. The shaded area represents the region of sensitivity with the upper binomial errors and amplitude uncertainties added to the variability fraction.} \label{sensitivity} \end{figure} \begin{figure} \centering \includegraphics[width=84mm]{Figure9.pdf} \caption{Variability frequency as a function of amplitude (dashed line) with the binomial errors and amplitude uncertainties added to the variability fraction (shaded area).} \label{var_fraction} \end{figure} \begin{figure} \centering \includegraphics[width=84mm]{Figure10.pdf} \caption{Percentage of simulated sinusoidal light curves detected as variable, as a function of period from 1~hour to 12~hours, for three different amplitudes. We used the measured survey median noise of 0.7\%, and each sine curve was sampled at intervals of 15 minutes to imitate the binned data of the survey. Additionally, we stepped through each sine curve at 5 degree phase intervals, to ensure that we sampled the full phase of the variable light curve.} \label{per_sensitivity} \end{figure} \begin{equation} B(n;N,{\epsilon}_v) = \frac{N!}{n!(N-n)!}{\epsilon}_v^n(1-{\epsilon}_v)^{N-n}. \label{equation3} \end{equation} \noindent where $n$ is the number of variables, $N$ the sample size and ${\epsilon}_v$ the variability frequency. This approach is based on Bayes' theorem under the assumption of a uniform prior based on no a priori knowledge and is ideal for small samples such as is the case for the BAM survey. The rotation period is another factor that can influence the detectability of a variable signal. In Figure~\ref{per_sensitivity}, we present the results of simulating light curves to test the detection probability of the survey to brown dwarf variables with different periods. We simulated sinusoidal light curves with three different amplitudes, of 1.5\%, 2.5\%, and 5.0\%, and with periods ranging from a minimum of 1~hour to a maximum of 12 hours \citep{zapatero06}. Gaussian noise equal to the median photometric uncertainty of the survey of 0.7\% was added to each light curve. To mimic the binned SofI data, the light curves were sampled at intervals of 15~minutes, and each simulated dataset was divided into groups of 3 hours, similar to the typical duration of the BAM data. For light curves with period longer than 3 hours, we generated multiple datasets, by stepping through the sine curve in steps of 5 degrees of phase and calculating the $p$-value at each phase, ensuring full sampling of the phase. Figure~\ref{per_sensitivity} shows the percentage of simulated light curves that are detected as variable with a $p$-value $\le5\%$. For amplitudes of 5.0\%, periodicities from 3 to 12~hours are easily recovered with a probability of 80 to 100\%, while the required periods decrease to $\sim$6~hours for 2.5\% variables and $\sim$5~hours for 1.5\% variables for detection probabilities in the 80 to 100\% range. \subsection{Frequency and amplitude of variability across spectral types} The frequency of variables as a function of spectral type is an important topic, since models of brown dwarf atmospheres have suggested that breakup of clouds across the L/T transition may result in both a higher rate of occurrence and a higher amplitude of variability compared to earlier L and later T objects. Amongst the previously known variables, the two largest amplitude variable objects discovered to-date are L/T transition objects - SIMP0136 ($\sim5$\% in $J$-band but with a significant night to night evolution, \citealt{artigau09}) and 2M2139 (as high as 26\% in the $J$-band, \citealt{radigan12}). \begin{table*} \centering \caption{Variability frequency.} \label{tabl:AmpLimits} \begin{tabular}{ccccc} \hline Sample & Sp. Type & No.~Targets & No.~Variables$\dagger$ & Freq. (\%)\\ \hline \hline Early-L & L0-L6 & 23 & 7 & $30^{+11}_{-8}$\\ Late-T & T5-T8 & 23 & 3 & $13^{+10}_{-4}$\\ L/T Transition & L7-T4 & 23 & 3 & $13^{+10}_{-4}$\\ Outside L/T transition & L0-L6 \& T5-T8 & 46 & 10 & $22^{+7}_{-5}$\\ \hline \end{tabular} \\Notes: $^\dagger$ These are the variables with a $p$-value~$\leq~5\%$, and amplitude $\ge2.3\%$. \end{table*} \begin{figure*} \centering \includegraphics[width=96mm]{Figure11a.pdf} \includegraphics[width=87mm]{Figure11b.pdf} \caption{Diagram on the left shows the amplitude of the variables ($p$-value~$\leq~5\%$ -- closed circles and $5\% <$ $p$-value~$\leq~10\%$ -- closed circles with cross) as well as the target photometric uncertainty of the non-varying objects (coloured triangles) across the entire spectral range of the sample. The diagram on the right shows the colour-colour diagram of the entire L through T spectral range with the full sample plotted with open circles, showing the colour spread of the targets. The variables from the BAM sample are overplotted ($p$-value~$\leq~5\%$ -- closed circles and $5\% <$ $p$-value~$\leq~10\%$ -- closed circles with cross). The L/T transition is indicated by the dashed lines} \label{survey} \end{figure*} As indicated by the variability frequencies reported in Table~\ref{tabl:AmpLimits}, the BAM results show no evidence that the frequency of variables in the L7 to T4 transition region is distinct from the earlier spectral types, the later spectral types, or the combination of all non-transition region brown dwarfs. The variability frequencies in Table~\ref{tabl:AmpLimits} are calculated using the entire sample of targets and an amplitude threshold of $\ge2.3\%$ and $p\le0.05$. Although no statistically significant difference in the variability frequencies for transition brown dwarfs is measured with the BAM survey, the 2.3\% amplitude limit of the analysis would not have detected differences at lower amplitudes, and removing the peak-to-trough amplitude threshold does not change this result. Adjusting the boundaries of the transition region by up to two spectral types does not change the result. Likewise, the amplitudes of the detected variables show no clear trends with spectral type within the capacity of the survey, as shown in Figure~\ref{survey} (left). The BAM variability frequency is very comparable to estimates for M stars. A variability frequency between $\sim$21--29\% for 19 M-stars was measured in a multi-wavelgenth optical study with the Calar Alto Observatory in Spain \citep{rockenfeller06}. The wavelength of observations for the M-star study was shorter than the BAM survey $J$-band data, and the amplitude of variations is expected to decline for longer wavelengths \citep[e.g.][]{reiners10}. In a recent compilation of variability surveys, \cite{khandrika13} reported a variability frequency of $30\pm5$\% based on a collection of different surveys with observations obtained in the optical and near-IR passbands, covering 78 objects in total. Comparison between surveys is difficult as the variability frequency may depend on a variety of different factors, including the target selection criterion and the criteria used to define variability in the targets which usually differs from one survey to the next. Additionally, the observed wavelength may also alter the variability frequency with different wavelength probing different depths in the atmosphere. \cite{koen13} finds a poor overlap between the variables identified with optical and near-IR filters (of the 13 variables already observed in near-IR surveys, 7 were found as constant and 6 as variable in the optical). Because of the uniform sensitivity of this survey, we did not incorporate the results of previous studies into the statistics, presented in Table~\ref{tabl:AmpLimits}. The presence of highly variable objects outside the transition region, may suggest the possibility of both early onset of cloud condensation in the atmospheres of mid-L dwarfs and the emergence of sulfide clouds in mid-T dwarfs \citep{Morley12}. Other physical processes that have been suggested to possibly induce variability in the atmospheres of brown dwarfs include coupling clouds with global atmosphere circulation \citep{showman13,zhang14}, and variability caused by thermal perturbations emitted from deeper layers within the brown dwarf atmosphere \citep{robinson14}. \subsection{Variability as a function of colour within a spectral type} The $J-K$ colour of the sample as a function of the spectral type is shown in Figure~\ref{survey} (right). The targets span nearly the full colour spread in early-L, transition and late-T sub sample. The 14 BAM variables and the three candidates are not clustered toward either the red or the blue within any particular spectral type. Previous studies \citep[e.g.][]{khandrika13} have suggested that brown dwarfs with unusual colours (highly red or blue) compared to the median of the spectral type might be indicative of variable cloud cover. We performed a two sample K-S test to determine whether or not the detrended colors of the BAM variables were distinct from the rest of the sample. The maximum difference between the cumulative distributions was 0.18 with a corresponding $p$-value of $\sim75\%$, indicating that the two datasets are consistent with being drawn from the same sample. The BAM study thus finds no correlation between the variables and the colour of a brown dwarf within each spectral type. \subsection{Binarity and variability} The BAM sample includes 12 confirmed binaries out of 47 targets studied for binarity with another four SpeX spectra binary candidates. Including the binary candidates, 10 out of the 16 binaries in the BAM sample fall in the L/T transition. This is consistent with previous detections of an increase in the binary frequency across the L-T transition \cite{burgasser06}. Amongst the BAM variables, only 2M2255 is a confirmed binary, while 2M1207 and 2M2139 are binary candidates. Five of the variables are confirmed to be single, and six have not been studied for binarity. The limited data provides no evidence to support a correlation between variability and binarity amongst the objects in the BAM survey. \subsection{Persistence of variability} \label{Persistence} A recent multi-epoch ($\sim$4~years) monitoring study of the variable brown dwarf SIMP0136 \citep{metchev13} revealed that the target has significant evolution in its light curve, changing from highly variable to constant in a 2~month period. When compared to the SofI light curve for SIMP0136, the target shows a fascinating variation in amplitude. It appears to be variable at 3\% in the SofI data, while a month later it shows large amplitude variations ($\sim$9\%) in the $J$-band, only to appear constant a few months later. Similar night-to-night variations have also been seen in SDSS J105213.51+442255.7 \citep{girardin13}. The evolution indicates a lack of persistence in the source of variability over timescales longer than a few weeks and it suggests that the brown dwarfs identified as constant in this study might similarly exhibit periods of quiescence and enhanced activity. The BAM survey only examines variability on the timescale of a single rotation period or less as compared to some surveys \citep[e.g.][]{gelino02,enoch03} that study the flux variations of brown dwarfs over longer timescales. The BAM data, in combination with previous results, can be used to address the question of persistence of variability. Table~\ref{tabl:PersRes} summarizes the observations related to persistence of variability, using information presented in Table~\ref{tabl:variable_sample} and \ref{tabl:LitConst}. For greatest consistency with the BAM study, we consider other epochs of near-IR data rather than optical. A total of of 34 BAM targets have an earlier epoch of observation. 2M0228 is the only source measured to be variable in the optical ($I_c$) that switched from variable to constant. Table~\ref{tabl:PersRes} indicates that brown dwarf variability does not necessarily persist on longer timescales, with only half the BAM variables showing variation in both epochs. The survey finds four previously constant objects to be variable and nine targets previously reported as variable in the literature to be constant, making these ideal candidates for multiple epoch monitoring programs. \begin{table} \centering \caption{Summary of Persistence Results} \label{tabl:PersRes} \begin{tabular}{lc} \hline Total targets with 2 epochs & 34 \\ \hline Variable at 2 epochs & 6 \\ Constant at 2 epochs & 15 \\ Switch between variable and constant & 13 \\ \hline \end{tabular} \end{table} | 14 | 4 | 1404.4633 |
1404 | 1404.6894_arXiv.txt | We discuss the $\gamma$--ray emission of the radiogalaxy NGC 1275 (the central galaxy of the Perseus Cluster), detected by {\it Fermi}--LAT and MAGIC, in the framework of the ``spine-layer" scenario, in which the jet is assumed to be characterized by a velocity structure, with a fast spine surrounded by a slower layer. The existence of such a structure in the parsec scale jet of NGC 1275 has been recently proved through VLBI observations. We discuss the constraints that the observed spectral energy distribution imposes to the parameters and we present three alternative models, corresponding to three different choices of the angles between the jet and the line of sight ($\theta_{\rm v}=6^{\circ}, 18^{\circ}$ and 25$^{\circ}$). While for the the case with $\theta_{\rm v}=6^{\circ}$ we obtain an excellent fit, we consider this solution unlikely, since such small angles seems to be excluded by radio observations of the large-scale jet. For $\theta_{\rm v}=25^{\circ}$ the required large intrinsic luminosity of the soft (IR--optical) component of the spine determines a large optical depth for $\gamma$--rays through the pair production scattering $\gamma \gamma\rightarrow e^+ e^-$, implying a narrow cut--off at $\sim50$~GeV. We conclude that intermediate angles are required. In this case the low frequency and the high--energy emissions are produced by two separate regions and, in principle, a full variety of correlations is expected. The correlation observed between the optical and the $\gamma$--ray flux, close to linearity, is likely linked to variations of the emissivity of the spine. | The extragalactic sky at very--high energies (VHE; $E>50$ GeV) is dominated by blazars, radio--loud active galactic nuclei with relativistic jets pointing toward the Earth. This geometry is particularly favorable since, due to the relativistic aberration, the resulting non--thermal emission of the jet is strongly amplified and blue shifted. The spectral energy distribution (SED) of blazars displays two broad components, the one peaking in the IR--UV band produced by relativistic electrons through synchrotron emission, and the high--energy one (peaking in $\gamma$--rays) attributed, in the so--called one--zone leptonic models, to the inverse Compton (IC) emission by the same electrons (e.g., Ghisellini et al. 1998). The parameter regulating the amplification of the emitted radiation, the relativistic Doppler factor $\delta\equiv[\Gamma(1-\beta\cos\theta _{\rm v})]^{-1}$ (where $\Gamma$ is the bulk Lorentz factor of the flow and $\beta=v/c$), is strongly dependent on the angle between the jet axis and the line of sight, $\theta_{\rm v}$: it is maximal within a cone with semi--aperture $\theta _{\rm v}\simeq 1/\Gamma$ and drops rapidly outside it. In this scheme thus one expects that for viewing angles larger than $1/\Gamma$ the jet non--thermal luminosity becomes increasingly fainter. Accordingly to this expectation, only a handful of radiogalaxies -- thought to be the {\it parent} (i.e. not beamed at us) population of blazars -- have been detected in VHE $\gamma$--rays (Aharonian et al. 2006, 2009, Aleksic et al. 2012). In the simple scheme sketched above, the SED of radiogalaxies should resemble that of blazars, the only difference being the reduced beaming. However, as discussed in e.g., Tavecchio \& Ghisellini (2008, TG08 hereafter) and Aleksic et al. (2014, A14 hereafter), the relatively large separation of the frequencies of two peaks, coupled to the expected low Doppler factor ($\delta\simeq 2-4$) is difficult to reproduce with one--zone emission models. Indeed, there are indications that the jet structure is not as simple as that assumed above. There is compelling observational evidence that the jet of low--power TeV blazars could be structured (e.g. Giroletti 2008), with a fast (with bulk Lorentz factor $\Gamma_{\rm s}=10-20$) central spine, surrounded by a slower ($\Gamma_{\rm l}=2-4$) layer. A structure of this type is also required to unify BL Lacs objects with the parent population of FRI radiogalaxies (Chiaberge et al. 2000). From the point of view of the radiative properties, such a structure leads to the enhancement of the radiative efficiency of both components, since the electrons of each region can inverse--Compton scatter the beamed soft photons coming from the other (Ghisellini et al. 2005). As shown in Ghisellini et al. (2005) this model can account for the VHE emission of TeV detected blazars (thought to be produced in the spine) and predicts that the weakly beamed emission of the layer could dominate the emission from misaligned radiogalaxies. This scheme was successfully applied to the modeling of the SED of the first radiogalaxy detected at VHE, M87 (TG08). The same idea is at the base of the decelerating jet scenario, in which one assumes that the inner jet moves faster then the outer jet zones. Also in this scenario the faster and the slower regions can interact through their radiation fields (Georganopoulos \& Kazanas 2003). NGC 1275, located in the Perseus cluster ($D\simeq 75$ Mpc) is one of the closest radiogalaxies. It has been the subject of intense study in the last years, particularly in view of the evident impact of the relativistic jets on the gas of the surrounding cluster (e.g. Fabian et al. 2011). At high energy it was not detected by EGRET onboard {\it CGRO} flown in the nineties. Instead, {\it Fermi--}LAT detected NGC 1275 soon after its launch (Abdo et al. 2009), with a flux 4 times larger than the EGRET upper limit, implying secular variability of the high-energy flux. The $\gamma$--ray emission is variable also on much smaller timescales, down to a week (Kataoka et al. 2010, Brown \& Adams 2011). Radio observations revealed an increasing of the activity starting in 2005 (Abdo et al. 2009) and recent VLBI observations (Nagai et al. 2010, 2014) show a new radio component ejected from the core. NGC 1275 has been the third radiogalaxy detected in the VHE band (by MAGIC), showing a very soft spectrum smoothly connected to the GeV spectrum (Aleksic et al. 2012). A14 report multifrequency data obtained during two multifrequency campaigns held in 2010--2011, thanks to which the overall SED could be assembled. Unfortunately, the available X-ray spectrum, probing the critical region of the SED between the two bumps, is affected by strong pile-up and it was not possible to fix the slope of the underlying non--thermal power law. The SED is barely compatible with a one-zone leptonic model for the values of the viewing angle usually considered in literature. The problem with the one--zone model is that, with the small Doppler factors corresponding to the large $\theta_{\rm v}$, it is difficult to reproduce the required large separation of the frequencies of the two peaks. Similar to the case of M87, it is tempting to think that a solution is to admit a structured jet. In fact, strong observational support to this scenario is provided by the clear evidence for a spine--sheath structure shown in recent VLBA maps presented in Nagai et al. (2014). The observed limb--brightened structure suggests the presence of a faster central core and a slower sheath for the inner (i.e. parsec scale) jet. For an assumed angle of $25^{\rm o}$ and with the hypothesis of equal rest-frame emissivity, the data are consistent with a bulk Lorentz flow of the sheath of 2.4. Motivated by this observational evidence and by the difficulties of the one-zone model, in the following we will explore the applicability of the spine--layer scenario of Ghisellini et al. (2005) to the multifrequency emission of NGC 1275. After a discussion of the SED (\S2) we will describe the model and its application (\S3). We discuss the results in \S4. Throughout the paper, we assume the following cosmological parameters: $H_0=70$ km s$^{-1}$ Mpc$^{-1}$, $\Omega_{\rm M}=0.3$, $\Omega_{\Lambda}=0.7$. | Summarizing, we have shown that the overall SED of NGC 1275 can be satisfactorily reproduced in the framework of the ``spine-layer" model --- with bulk Lorentz factors for the two components similar to those used in the case of M87 --- if the viewing angle is smaller then $\theta_{\rm v}\sim 20$ deg. Larger angles inevitably lead to a drastic suppression of the emission in the MAGIC energy band caused by the strong internal absorption of the $\gamma$--rays with energies above few tens of GeV, determined by the luminous IR radiation field associated to the spine emission. Estimates of the angle at pc scale, based on the detection of the counter jet, provide values in the range 30--55 degrees (Walker et al. 1994, Asada et al. 2006). However, smaller angles are often assumed (e.g. Abdo et al. 2009). A much lower value ($\theta_{\rm v}\approx 3^{\rm o}$) at sub--pc scale was derived from VLBI observations by Krichbaum et al. (1992). Considering these uncertainties and the not unlikely possibility that the jet bends from the sub--pc scale (where we suppose that the emission occurs) to the pc scales imaged with VLBI, we conclude that the spine--layer scenario is clearly constrained but still barely suitable to reproduce the data. A strong test able to seriously threaten the model would be the detection of photons at energies above 1 TeV, for which, even with $\theta_{\rm v}=18^{\rm o}$, the optical depth $\tau_{\gamma \gamma}>10^3$, corresponding to a suppression of the flux exceeding $10^3$. A point not completely satisfactory about our preferred model for $\theta_{\rm v}=18^{\rm o}$ concerns the emission the jet would present if observed at smaller angles. As previously discussed in TG08, in the unification scheme for radio--galaxies and blazars we generally expect that, once observed at small angles, the non--thermal continuum from a radiogalaxy resembles that of a blazar. In particular, for NGC 1275 we expect that the SED shape follows that of high--power BL Lac objects. We indeed found a source whose SED shape closely traces our theoretical curve, but with a luminosity about one order of magnitude smaller than the beamed emission of NGC 1275. A14 report the existence of a clear correlation between the LAT $\gamma$--ray flux, $F_{\gamma}$, and the optical flux, $F_{\rm opt}$, compatible with both a linear ($F_{\gamma}\propto F_{\rm opt}$) and a quadratic ($F_{\gamma}\propto F_{\rm opt}^2$) dependence. In our scheme the flux in the two bands is dominated by the emission produced in two separate regions, i.e. the spine for the optical and the layer for the $\gamma$--rays. In principle several possibilities are allowed by this configuration. For instance, $\gamma$--ray variability without counterpart in the optical is possible with changes of the parameters of the layer (in particular those specifying the electron energy distribution) and a stationary spine emission. On the other hand, since the IC emission of the layer --- responsible for the observed high--energy emission --- is dominated by the scattering of the synchrotron photons of the spine, we expect that variations of the optical emission are accompanied by variations in the $\gamma$--ray band. If such variations are driven by changes in the intrinsic emissivity of the spine, without changes in the structural parameters (in particular the bulk Lorentz factor) we expect a linear correlation. However, if variations of $\Gamma_{\rm s}$ (and/or of $\Gamma_{\rm l}$) are also involved, more complex patterns are possible, connected to the interplay between the beaming of the emission in the observer frame and the amplification of the energy density of the seed photons, depending on the relative motion of the spine and the layer. With such a complex phenomenology is difficult to clearly identify a mechanism for the observed correlation. The simplest situation, corresponding to a linear correlation, is associated to variations of the electron energy distribution or of the magnetic field of the spine. | 14 | 4 | 1404.6894 |
1404 | 1404.4069_arXiv.txt | We explore the hypothesis that a passing satellite or dark matter subhalo has excited coherent oscillations of the Milky Way's stellar disk in the direction perpendicular to the Galactic midplane. This work is motivated by recent observations of spatially dependent bulk vertical motions within $\sim 2$ kpc of the Sun. A satellite can transfer a fraction of its orbital energy to the disk stars as it plunges through the Galactic midplane thereby heating and thickening the disk. Bulk motions arise during the early stages of such an event when the disk is still in an unrelaxed state. We present simple toy-model calculations and simulations of disk-satellite interactions, which show that the response of the disk depends on the relative velocity of the satellite. When the component of the satellite's velocity perpendicular to the disk is small compared with that of the stars, the perturbation is predominantly a bending mode. Conversely, breathing and higher order modes are excited when the vertical velocity of the satellite is larger than that of the stars. We argue that the compression and rarefaction motions seen in three different surveys are in fact breathing mode perturbations of the Galactic disk. | Recently, three independent surveys of stellar kinematics within $\sim 2\,{\rm kpc}$ of the Sun detected spatially dependent bulk motions in the direction perpendicular to the Galactic plane \citep{widrow2012, williams2013, carlin2013}. \citet{widrow2012} found that the bulk motions of stars, when plotted as a function of position $z$ relative to the Galactic midplane, have characteristics of a breathing mode perturbation with a velocity gradient of $\sim 3-5\,{\rm km\,s}^{-1}\,{\rm kpc}^{-1}$. This result was based on a sample of 11K main sequence stars from the Sloan Extension for Galactic Understanding and Exploration (SEGUE) survey, which focused on intermediate latitudes above and below the Galactic midplane and Galactic longitudes in the range $100^\circ < l < 180^\circ$ \citep{yanny2009}. \citet{williams2013} used a sample of 72K red-clump stars from the Radial Velocity Experiment (RAVE) survey \citep{steinmetz2006} to map out bulk motions as a function of Galactocentric radius and $z$ and found evidence for compressional motion outside the solar circle and rarefaction inside with peak vertical bulk velocities of $\pm 15\,{\rm km\,s}^{-1}$. \citet{carlin2013} found similar features in their analysis of 400K F-type stars with proper motions from the PPMXL catalog \citep{roeser2010} and spectroscopic radial velocities from the LAMOST/LEGUE survey \citep{cui2012, zhao2012}. As stressed by \citet{carlin2013}, the three surveys look in different parts of the extended solar neighborhood and consider different spatial projections of the data. \citet{widrow2012} also found a North-South asymmetry in the number counts when plotted against $z$. The number count asymmetry was confirmed by \citet{yanny2013} who carried out a careful analysis of the uncertainties and potential systematic effects. The number counts show a 10\% (North - South)/(North + South) deficit at $|z|\simeq 400 {\rm pc}$ and an excess of about the same magnitude at $|z|\simeq 800 \,{\rm pc}$. It is possible that the North-South asymmetries in number counts and bulk vertical motions are the result of stellar debris from a tidally disrupted satellite galaxy that is mixing in with the disk stars. \citet{widrow2012} considered an alternative hypothesis in which the North-South asymmetries arise from coherent oscillations of the disk itself, which were excited by a passing satellite or dark matter subhalo. Eventually, the oscillations die away due to phase mixing and Landau damping. The disk settles into a new equilibrium state, albeit one with a higher velocity dispersion. Thus, the bulk motions seen in the data may indicate an early phase of a disk heating event. The aim of this paper is to explore this hypothesis in more detail through toy model calculations and numerical simulations. A similar question was posed by \citet{minchev2009} in the context of the stellar velocity distribution in the solar neighborhood. Detailed analyses of {\it Hipparcos} data \citep{chereul1998, dehnen1998, chereul1999,nordstrom2004} have revealed rich substructure in the local velocity distribution. There are classical moving groups, which are thought to be the stellar streams from dissolved star clusters (see \citet{eggen1996} and references therein). Velocity-space features may arise from dynamical effects of the bar \citep{dehnen2000} or spiral structure \citep{desimone2004,quillen2005,chakrabarty2007} or they may be due to streams of stars that were tidally stripped from accreted satellite galaxies \citep{navarro2004, helmi2006}. \citet{minchev2009} consider a different scenario in which the ``energy kick'' from a passing satellite leaves ripples in the (disk-plane) velocity distribution of stars. Their conjecture is that velocity-space substructure in the solar neighborhood is a manifestation of these ripples. Satellites, dark matter subhalos, and globular clusters for that matter, have long been recognized as possible culprits of disk heating and thickening. In general, a massive object that passes through the disk will transfer a fraction of its orbital energy to the disk stars \citep{lacey1985,toth1992, sellwood1998}. Satellite interactions can also cause the disk to spread out radially and develop warps and flares \citep{quinn1986,quinn1993,walker1996,velazquez1999} Satellite encounters can excite various modes in the disk such as bending modes and breathing modes \citep{toomre1966,araki1985,mathur1990,weinberg1991}. It is the latter that corresponds most closely to the velocity perturbations seen in the SEGUE, RAVE, and LAMOST surveys. We will show that a bending mode perturbation arises when the satellite's vertical velocity is less than that of the disk stars while breathing and higher-order modes are excited when the vertical velocity of the satellite exceeds that of the stars. Satellites and subhalos can also excite spiral structure and bars in stellar disks. (See \citet{sellwood2013} for a recent review of the stellar dynamics of disk galaxies.) The seminal work of \citet{toomre1972} showed that the tidal interaction between a stellar disk and a companion galaxy of comparable mass can generate grand design spiral structure similar to what is seen in M51. Alternatively, multi-armed and flocculent spiral structure can arise from the continual interactions between the disk and a system of satellite galaxies and dark matter subhalos. Cosmological simulations of structure formation in a $\Lambda$CDM universe suggest that the halos of Milky Way size galaxies harbour a wealth of substructure in the form of subhalos \citep{klypin1999,moore1999,gao2004}. These results motivated \citet{gauthier2006} and \citet{dubinski2008} to explore satellite-disk interactions for an M31-like galaxy. For their particular M31 model, when the halo is smooth the disk remains stable against bar formation for 10 Gyr and relatively weak spiral structure develops, presumably seeded by the shot noise of the N-body realization. Conversely, when $\sim 10\%$ of the halo mass initially resides in compact subhalos, the disk develops prominent spiral features and forms a strong bar. Similar results were found in a series of simulations by \citet{kazantzidis2008}. More recently, \citet{purcell2011} considered a model Milky Way disk that was perturbed by a single satellite galaxy. The prototype for their perturber was the Sagittarius dwarf spheroidal galaxy \citep{ibata1994, ibata1997}, which is believed to have survived several orbits about the Galaxy. In the \citet{purcell2011} simulations spiral structure emerges that is similar to the spiral structure observed in the Milky Way. \citet{gomez2013} reanalysed these simulations and found that vertical perturbations in the number density of disk particles at roughly the Sun's position from the Galaxy's center were also generated and qualitatively similar to those seen in \citet{widrow2012} and \citet{yanny2013}. It is not at all surprising that a Sagittarius-like dwarf produces vertical oscillations similar to what is found in the data. The solar neighborhood is characterized by a circular speed about the Galactic center of $\simeq 220-230\,{\rm km\,s}^{-1}$ (see, for example, \citet{bovy2012c} and references therein) and a stellar surface density of $\simeq 50\,M_\odot\,{\rm pc}^{-2}$ (see, for example, \citet{holmberg2004}). Stars in the disk have vertical velocities in the range of $10-40\,{\rm km\,s}^{-1}$ \citep{robin2003,bovy2012a}. A satellite with a comparable surface density to that of the solar neighborhood, with an orbit that is matched to the local standard of rest (LSR), and with a vertical velocity through the midplane in resonance with the vertical motions of disk stars will produce the strongest perturbations. In Section 2, we review earlier discussions of disk heating and thickening and then present a simple toy-model calculation for the excitation of bending and breathing modes. In Section 3, we present results from one-dimensional N-body simulations that support the toy-model calculation and also illustrate the potentially long-lived nature of the oscillations. We provide preliminary results from fully self-consistent 3D N-body simulations of satellite-disk encounters in Section 4. Finally, we summarize our results and give concluding remarks in Section 5. | The implications of spatially dependent bulk motions perpendicular to the Galactic disk were highlighted by \citet{oort1932} in his seminal work on the structure of the Galactic disk. Oort's aim was to determine the potential $\psi(z)$ from the local stellar density and velocity distribution. He based his analysis on the assumption that the local distribution of stars is in equilibrium. To test the assumption, he computed the mean vertical velocity for stars in four separate bins: $100\,{\rm pc} < \pm z < 200\,{\rm pc}$ and $200\,{\rm pc} < \pm z < 500\,{\rm pc}$ but did not find evidence for systematic motions, a result that he notes ``lends some support to the assumption \ldots that in the $z$-direction the stars are thoroughly mixed'' \citep{oort1932}. Turning Oort's argument around, the detection of bulk vertical motions by the SDSS/SEGUE, RAVE, and LAMOST surveys suggests that the local Galactic disk is not in equilibrium in the $z$-direction. In this paper, we considered the hypothesis that the observed bulk vertical motions were generated by a passing satellite or dark matter subhalo. The idea that dark matter, in one form or another, might be responsible for heating and thickening the disk dates back to the 1980's. \citet{lacey1985} calculated disk heating by a dark halo of supermassive black holes while \citet{carr1987} considered dark matter in the form of $10^6\,M_\odot$ dark clusters. In essence, our hypothesis is that the bulk motions seen in the data represent the early stages of a disk heating event. Our focus has been to explore the theoretical aspects of disk-satellite interactions. We found that the nature of the perturbations is controlled largely by the satellite's vertical velocity relative to the disk. In particular, a slow moving (as measured in the LSR) satellite induces a bending mode perturbation. With a higher vertical velocity, higher order modes, such as the breathing mode, are excited. Thus, if a satellite is indeed responsible for the bulk vertical motions seen in the solar neighborhood, this its vertical velocity through the disk would likely have been $\appgeq 50\,{\rm km\,s}^{-1}$. Moreover, its surface density would have to be comparable to that of the disk in order to produce an appreciable perturbation. The model satellites considered by \citet{gomez2013} satisfy these conditions and so it is not surprising that they found vertical perturbations in the disk that were qualitatively similar to what was found in the data. Single satellite simulations show that after a localized breathing mode perturbation is produced, it is sheared by the differential rotation of the disk. After several orbital periods of the disk, the perturbation assumes on a spiral-like pattern. The situation is more complicated with a population of satellites and it may be difficult to disentangle initial perturbations from the accumulated long-lived perturbations. Our analysis, and that of \citet{weinberg1991} suggest that stars on the tail of the energy distribution are most responsive to a breathing mode perturbation. Though our analyses focused on single-component disks, it may well be that the vertical motions is a property more of the thick disk stars, than the thin disk stars. It is well known that the vertical velocity dispersion and scale height are anti-correlated with metallicity (see, for example, \citet{bovy2012a,bovy2012b,minchev2013,minchev2014} and references therein). The arguments presented in this paper suggest that bulk motions should be more prominent in the low metallicity/high $E_z$ populations. A related issue is radial migration. \citet{sellwood2002} argued that spiral waves can change the angular momenta, and hence Galactocentric radii, of individual stars by $\sim $50\%. In analysing the cosmological simulations of disk formation by \citet{martig2012}, \citet{minchev2013} found that satellite interactions can also drive radial migration. Moreover, radial migration can bring high dispersion stars from the inner disk to the solar neighborhood, and, as discussed above, these are the stars most susceptible to a recent satellite interaction. There are twenty-five known satellites of the Milky Way. Moreover, in a $\Lambda$CDM cosmology, the dark halo of a Milky Way-size galaxy is expected to harbour many more nonluminous subhalos \citep{klypin1999,moore1999}. Thus, it is likely that the Galactic disk has been continually perturbed over its lifetime. In principle, observations of bulk motions in the stellar disk could provide a probe of the subhalo distribution. To do so will require a suite of simulations where the slope and amplitude of the subhalo mass function are varied. Over the next few years, Gaia will provide an unprecedented snapshot of the Galaxy by making astrometric, spectral, and photometric observations of approximately one billion Milky Way stars (See, for example \citet{perryman2001} and \citet{bruijne2012}). This data set will yield a more accurate and complete map of bulk motions in the stellar disk. By bringing together these observations, theoretical analysis, and N-body simulations we hope to better understand Galactic dynamics, and in particular, interactions between the Milky Way's disk and its satellites and dark matter subhalos. | 14 | 4 | 1404.4069 |
1404 | 1404.4981_arXiv.txt | We have used optical observations of resolved stars from the Panchromatic Hubble Andromeda Treasury (PHAT) to measure the recent ($< 500\,\mathrm{Myr}$) star formation histories (SFHs) of 33 FUV-bright regions in M31. The region areas ranged from $\sim 10^4$ to $10^6\,\mathrm{pc}^2$, which allowed us to test the reliability of FUV flux as a tracer of recent star formation on sub-kpc scales. The star formation rates (SFRs) derived from the extinction-corrected observed FUV fluxes were, on average, consistent with the 100-Myr mean SFRs of the SFHs to within the $1-\sigma$ scatter. Overall, the scatter was larger than the uncertainties in the SFRs and particularly evident among the smallest regions. The scatter was consistent with an even combination of discrete sampling of the initial mass function and high variability in the SFHs. This result demonstrates the importance of satisfying both the full-IMF and the constant-SFR assumptions for obtaining precise SFR estimates from FUV flux. Assuming a robust FUV extinction correction, we estimate that a factor of 2.5 uncertainty can be expected in FUV-based SFRs for regions smaller than $10^5\,\mathrm{pc}^2$, or a few hundred pc. We also examined ages and masses derived from UV flux under the common assumption that the regions are simple stellar populations (SSPs). The SFHs showed that most of the regions are not SSPs, and the age and mass estimates were correspondingly discrepant from the SFHs. For those regions with SSP-like SFHs, we found mean discrepancies of $10\,\mathrm{Myr}$ in age and a factor of 3 to 4 in mass. It was not possible to distinguish the SSP-like regions from the others based on integrated FUV flux. | A common technique for estimating global star formation rates (SFRs) in individual galaxies is to measure the total flux at wavelengths known to trace recent star formation (SF), such as ultraviolet (UV) emission from intermediate- and high-mass stars. After correcting for dust extinction, an observed flux can be converted into a SFR using a suitable calibration, which is typically a linear scaling of intrinsic luminosity derived from population synthesis modeling. The modeling process requires a set of stellar evolution models and a stellar initial mass function (IMF), as well as a characterization of the star formation history (SFH; the evolution of SFR over time) and the metallicity of the population. These quantities are often not well-constrained for a given system and need to be assumed (see reviews by \citealt{Kennicutt98}, \citealt{Kennicutt12}, and references therein). A set of flux calibrations widely used in extragalactic studies were presented by \citet[][see \citealp{Kennicutt12} for updates]{Kennicutt98}. These calibrations are based on models of a generic population with solar metallicity, a fully populated IMF, and a SFR that has been constant over the lifetime of the tracer emission ($\sim 100\,\mathrm{Myr}$ for UV). The flux calibrations are therefore applicable to any population that can be assumed to approximate the generic population, such as spiral galaxies. In environments with low total SF (i.e., low mass) or on subgalactic scales, however, the assumptions of a fully populated IMF and a constant SFR start to become tenuous. As a result, applying the flux calibrations in these situations can lead to inaccurate SFR estimates. For populations located within a few Mpc, it is possible to measure SFRs more directly by fitting the color magnitude diagram (CMD) of the resolved stars to obtain a SFH \citep{Dolphin02}. At its core, CMD fitting is a population synthesis technique just like flux calibration (albeit much more complex) and thus requires a set of stellar evolution models, an IMF, and an accounting of dust. The primary advantage of CMD fitting over the flux calibration method for obtaining SFRs, however, is the elimination of assumptions about the SFH and metallicity. CMD-based SFHs thus provide a relative standard for testing the accuracy of SFR estimates from commonly used flux calibrations, especially in applications where the underlying full-IMF and constant-SFR assumptions are not strictly satisfied. More generally, the SFHs can be used to test results from any other flux-based method, such as ages and masses derived under the simple stellar population (SSP) assumption. With recent Hubble Space Telescope (HST) observations from the Panchromatic Hubble Andromeda Treasury \citep[PHAT;][]{Dalcanton12}, we have measured the recent SFHs ($< 500\,\mathrm{Myr}$) of 33 UV-bright regions in M31 and compared them with SFRs derived from UV flux. We also compared the SFHs with ages and masses derived from UV flux by treating the regions as SSPs. The UV-bright regions were cataloged by \citet[][\citetalias{Kang09} hereafter]{Kang09} using Galaxy Evolution Explorer (GALEX) far-UV (FUV) flux and have areas ranging from $10^4$ to $10^6\,\mathrm{pc}^2$. This range of sizes allows us to test the reliability of the full-IMF, constant-SFR, and SSP assumptions on sub-kpc scales. This paper is organized as follows. We describe our sample of UV-bright regions and show their CMDs from the PHAT photometry in \S \ref{observations}. We summarize the CMD-fitting process, describe our extinction model, and present the resulting SFHs of the regions in \S \ref{sfhs}. \S \ref{fluxmod} describes the modeling of UV magnitudes from the SFHs, and \S \ref{sfrs} describes the total masses and the mean SFRs from the SFHs, as well as the SFRs based on UV flux. In \S \ref{discussion}, we compare the UV flux-based SFRs, ages, and masses with the results from the SFHs, discuss the applicability of the full-IMF, constant-SFR, and SSP assumptions to our sample, and attempt to quantify the uncertainties associated with using UV flux to estimate SFRs, ages, and masses for sub-kpc UV-bright regions. \begin{figure} \centering \includegraphics[width=\columnwidth]{fig1.pdf} \caption{Two-color composite mosaic of M31 from the GALEX Deep Imaging Survey (FUV in blue, NUV in orange). The HST/ACS outlines of the PHAT survey area and Brick 15 are highlighted in blue and orange, respectively. Brick 15 covers a portion of the 10-kpc star-forming ring. The scale bar indicates a distance of $5\;\mathrm{kpc}$ along both the major and minor axes of M31 assuming an inclination of $78\;\mathrm{deg}$ \citep{Tully94}. } \label{fig:map_full} \end{figure} \begin{figure} \centering \includegraphics[width=\columnwidth]{fig2.pdf} \caption{Closeup of Brick 15 from the same image in Figure \ref{fig:map_full}. Brick 15 contains 33 of the UV-bright regions from the \citet{Kang09} catalog, highlighted in blue and labeled by ID number (see Table \ref{tab:observations}). The region areas, deprojected assuming an inclination of $78\;\mathrm{deg}$ \citep{Tully94}, range from $\sim 10^4$ to $10^6\;\mathrm{pc}^2$. The scale bar indicates a distance of $500\;\mathrm{pc}$ along both the major and minor axes of M31. } \label{fig:map_b15} \end{figure} | 14 | 4 | 1404.4981 |
|
1404 | 1404.4948_arXiv.txt | We summarize the nuclear physics interests in the Oklo natural nuclear reactors, focusing particularly on developments over the past two decades. Modeling of the reactors has become increasingly sophisticated, employing Monte Carlo simulations with realistic geometries and materials that can generate both the thermal and epithermal fractions. The water content and the temperatures of the reactors have been uncertain parameters. We discuss recent work pointing to lower temperatures than earlier assumed. Nuclear cross sections are input to all Oklo modeling and we discuss a parameter, the $^{175}$Lu ground state cross section for thermal neutron capture leading to the isomer $^{176\mathrm{m}}$ Lu, that warrants further investigation. Studies of the time dependence of dimensionless fundamental constants have been a driver for much of the recent work on Oklo. We critically review neutron resonance energy shifts and their dependence on the fine structure constant $\alpha$ and the ratio $X_q=m_q/\Lambda$ (where $m_q$ is the average of the $u$ and $d$ current quark masses and $\Lambda$ is the mass scale of quantum chromodynamics). We suggest a formula for the combined sensitivity to $\alpha$ and $X_q$ that exhibits the dependence on proton number $Z$ and mass number $A$, potentially allowing quantum electrodynamic and quantum chromodynamic effects to be disentangled if a broader range of isotopic abundance data becomes available. | On 25 September 1972, Andr\`{e} Giraud, Head of the French {\it Commissariat \`{a} l'\'{E}nergy Atomique} (CEA), announced the discovery of a two billion year-old nuclear reactor in Gabon, at the site of the Oklo uranium mines. The sequence of events that led to this startling announcement had begun earlier, in June 1972, at the Pierrelatte uranium enrichment plant with the observation of a small but definite anomaly in the uranium isotopic ratio for a UF$_6$ sample. The supply of anomalous uranium was soon traced to uranium rich ores (up to 60\%) for which investigations revealed local uraninite deposits with $^{235}$U isotopic abundance of 0.600\%, instead of the normal 0.7202\%. At first glance one can wonder how the Oklo reactors were able to operate when it is well known that the modern light water reactors cannot work with natural uranium, requiring instead $^{235}$U enrichment of about 3.5\%. Natural uranium is composed of three isotopes with abundances today\footnote{Abundance and lifetime data are from the National Nuclear Data Center (www.nndc.bnl.gov).} of 99.2744\% for $^{238}$U, 0.7202\% for $^{235}$U, and 0.0054\% for $^{234}$U. However, the relative enrichment of $^{235}$U increases going back in geological time because the half-life of $^{235}$U is 710~Myr while that of $^{238}$U is 4.51~Gyr. For example, $^{235}$U enrichment was 1.3\% 700~Myr ago, 2.3\% 1.40~Gyr ago, 4.0\% 2.10~Gyr ago, and up to 17\% at the time of creation of the solar system. As reviewed by Zetterstr\" {o}m \cite{Lena00}, the stabilization of the Oklo area geological basement happened not earlier than 2.7 Gyr ago and the geological age of the Francevillian sediments is estimated \cite{Gau96b} to be about $2.265\pm 0.15$~Gyr. An independent constraint on the age of the Oklo phenomenon is provided by the Great Oxidation Event \cite{GauW03,Haz09}. This happened about 2.2 Gyr ago when, due to the biological activity of cyanobacteria, the oxygen content in the atmosphere of Earth increased by about a factor of a hundred. This allowed uranium to be converted from its insoluble uranium(IV) form to its soluble uranium(VI) form. Deposition of high grade uranium ores in sediments subsequently occurred when this soluble uranium was precipitated out, either by reduction back to the insoluble form (by carbon, methane or iron), or by direct microbial induced action. Interest in natural nuclear reactors preceded their discovery by almost two decades. As early as 1953, Wetherhill and Ingram had found evidence in Congo pitchblende Xe isotopic data that, besides spontaneous fission, neutron induced fission had taken place \cite{WeIn53}. They stated that ``the deposit was twenty-five percent of the way to becoming a pile; it is interesting to extrapolate back 2000 million years \ldots\ Certainly such a deposit would be close to being an operating pile.'' Following this suggestion, while considering the Johanngeorgenstadt (Saxony) pitchblende -- a uranium ore with a minimal content of rare earth poisons, Kuroda \cite{Kur56} applied Fermi's four-factor pile theory \cite{Fer47} and obtained estimates of neutron multiplication factors greater than unity for proper amounts of water in pitchblende. His conclusion was that ``critical uranium chain reactions could have taken place if the size of the assemblage was greater than, say, a thickness of a few feet.'' Kuroda looked for possible changes in the chemical composition of samples from uranium ores from several locations (but not Oklo, of course), and found no signs of {\em chain} reactions. At the time his paper was written (1956), it seemed highly unlikely that natural reactors would be found on Earth. The first research reports on Oklo data appeared in the fall of 1972 (Bodu {\em{et al.}} \cite{Bodu72}, Neuilly {\em{et al.}} \cite{Neu72} and Baudin {\em{et al.}} \cite{Bau72}) and starting in 1973, the CEA launched the project ``Franceville'', named after the town Franceville in the vicinity the Oklo mines. Uranium mining was suspended for two years to probe the terrain. Six natural reactor zones were discovered and samples were shared widely with the cooperation of the International Atomic Energy Agency (IAEA). In June 1975, the first international Oklo meeting took place in Libreville, with the proceedings published by the IAEA \cite{IAEA75}. To continue the work, the IAEA and CEA established an International Working Group on Natural Reactors with a technical committee of experts. This group met in Paris in December 1977 to review progress, and published further proceedings in 1978 \cite{IAEA78}. A review of all work done until 1990 can be found in an excellent book by Naudet \cite{Nau91}, who was the Franceville project head. New zones were later identified \cite{Gau96a}, and a European research program ``Oklo - natural analogue for a radioactive waste repository'' was initiated to study analogies between the behavior of materials in Oklo and in planned nuclear waste repositories \cite{Bla96}. This program was financed by the European Commission of the European Union and implemented in co-operation with institutions from other countries. Studies of Oklo continue unabated with about 140 papers in the published literature since 2000. Recent papers split evenly between interest in the fascinating geology and operation of the reactors, and interest in what the isotopic remains can say about time variation of fundamental constants over the last two billion years. Notable earlier reviews include those by Naudet \cite{Nau76}, Cowan \cite{Cow76}, Petrov \cite{Pet77}, Kuroda \cite{Kur82}, Meshik \cite{Mesh05}, and Barr\`{e} \cite{Bar05}. We use the basic information from these reviews and focus here on recent results on modeling the reactors and their implications for refining bounds on the time variation of dimensionless fundamental constants such as the electromagnetic fine structure constant. | Unravelling how the geosphere and the biosphere evolved together is one of the most fascinating tasks for modern science. The Oklo natural nuclear reactors, basically formed by cyanobacteria two billion years ago, are yet another example of the surprises to be found in Earth's history. Since their discovery over forty years ago, the reactors have provided a rich source of information on topics as applied as can nuclear wastes be safely stored indefinitely to topics as esoteric as are the forces of physics changing as the Universe ages? In this review, we have summarized nuclear physics interests in the Oklo phenomenon, focusing particularly on developments over the past two decades. Modeling the reactors has become increasingly sophisticated, employing Monte Carlo simulations with realistic geometries and materials which can generate both the thermal and epithermal fractions. The water content and the temperatures of the reactors have been uncertain parameters. We have discussed recent work pointing to lower temperatures than earlier assumed. Nuclear cross sections are input to all Oklo modeling and we have identified a parameter, relating to the capture by the $^{175}$Lu ground state of thermal neutrons, that warrants further investigation. The use of Oklo data to constrain changes in fundamental constants over the last 2 billion years has motivated much recent work. We have presented a critical reappraisal of the current situation, starting with the long-standing study of Damour and Dyson on sensitivity to the fine structure constant $\alpha$. We conclude that their result can plausibly be used for order of magnitude estimates, but an investigation of how this conclusion may be affected by a more careful treatment of the Coulomb potential in the vicinity of the nuclear surface (and beyond) is warranted. The more recent analysis by Flambaum and Wiringa of the sensitivity to the average mass $m_q$ of the light quarks has been updated to incorporate the latest values of sigma terms. No firm conclusions about the reliability of Flambaum and Wiringa's estimate are possible (because of uncertainties surrounding the short-ranged part of the nuclear interaction), but it could be an overestimate by as much as an order of magnitude. On the basis of the work in Refs.~\cite{Dam96} and \cite{Flam09}, we have suggested a formula for the unified treatment of sensitivities to $\alpha$ and $m_q$, namely Eq.~(\ref{glb}). It is an obvious synthesis, which has the advantage of making explicit the dependence on mass number and atomic number. We hope that it may prove useful in distinguishing between the contributions of $\alpha$ and $m_q$ in a model independent way or, at least, facilitating an understanding of the significance of the non-zero shifts in resonance energies which have been found in some studies of Oklo data. Appealing to recent data on variations in $\alpha$ and the proton-to-electron mass ratio $\mu$, we have demonstrated that, within the very general model of Ref.~\cite{Lan02} (and contrary to widespread opinion), shifts in resonance energies are not any more sensitive to variations in $m_q$ than they are to variations in $\alpha$. When extracting an \emph{order of magnitude limit\/} on any change in $\alpha$, it is, thus, permissable to ignore any changes in $m_q$ and {\emph{vice versa}} (provided the model of Ref.~\cite{Lan02} applies). In fact, we have argued that one can, at best, use null shifts to establish order of magnitude estimates of upper bounds. Bounds on $\Delta\alpha/\alpha$ have been presented. The most recent study \cite{Oneg012} of the Oklo data pertaining to the $97.3\,\mathrm{meV}$ resonance seen now in neutron capture by ${}^{149}$Sm implies that $|\Delta\alpha/\alpha| \lesssim 1\times 10^{-8}$ (cf.~Table \ref{tb:alphabds} for more details). | 14 | 4 | 1404.4948 |
1404 | 1404.6660_arXiv.txt | We consider the possible detection of parity violation at the linear level in gravity using polarized anisotropies of the cosmic microwave background. Since such a parity violation would lead to non-zero $TB$ and $EB$ correlations, this makes those {\it odd-parity} angular power spectra a potential probe of parity violation in the gravitational sector. These spectra are modeled incorporating the impact of lensing and we explore their possible detection in the context of small-scale (balloon-borne or ground-based) experiments and a future satellite mission dedicated to $B$-mode detection. We assess the statistical uncertainties on their reconstruction using mode-counting and a (more realistic) pure pseudospectrum estimator approach. Those uncertainties are then translated into constraints on the level of parity asymmetry. We found that detecting chiral gravity is impossible for ongoing small-scale experiments. However, for a satellite-like mission, a parity asymmetry of 50\% could be detected at 68\% of confidence level (at least, depending on the value of the tensor-to-scalar ratio), and a parity asymmetry of 100\% is measurable with {\it at least} a confidence level of 95\%. We also assess the impact of a possible miscalibration of the orientation of the polarized detectors, leading to spurious $TB$ and $EB$ cross-correlations. We show that in the context of pseudospectrum estimation of the angular power spectra, self-calibration of this angle could significantly reduce the statistical significance of the measured level of parity asymmetry (by {\it e.g.} a factor $\sim2.4$ for a miscalibration angle of 1 degree). For chiral gravity and assuming a satellite mission dedicated to primordial $B$-mode, a non detection of the $TB$ and $EB$ correlation would translate into an upper bound on parity violation of 39\% at 95\% confidence level for a tensor-to-scalar ratio of 0.2, excluding values of the (imaginary) Barbero-Immirzi parameter comprised between 0.2 and 4.9 at 95\% CL. | \label{sec:intro} The anisotropies of the cosmic microwave background (CMB) are currently the most powerful probe of the physics underlying the primordial universe. Those anisotropies arise into three flavors: total intensity and two degrees of freedom describing its linear polarization. Though CMB polarized anisotropies are measured using two Stokes parameters, $Q$ and $U$, they are most conveniently described using a gradient-like, $E$-mode, and a curl-like, $B$-mode \cite{zaldarriaga_seljak_1997,kamionkowski_etal_1997}. Such a decomposition of the linearly polarized anisotropies is meaningful at a physical level as it is directly linked to the primordial cosmological perturbations sourcing CMB anisotropies. For instance, on the linear level the $B$-modes can be sourced by the primordial gravitational waves \cite{seljak_zaldarriaga_1997,spergel_zaldarriaga_1997} and not by the scalar fluctuations, thought to be largely responsible for the observed total intensity and $E$-mode anisotropies. Consequently, a detection of the $B$-mode anisotropy at large angular scales ($\ell \apprle 100$) in excess of what is expected from the gravitational lensing signal could be seen as a direct validation of inflationary theories, as the latter are considered to be the most likely source of the gravity waves, and could allow for discrimination between different inflationary models. It could also set useful constraints on the reionization period \cite{zaldarriaga_1997}. At smaller angular scales, $B$-modes are expected to be mainly due to gravitational lensing of CMB photons which converts $E$-modes into $B$-modes \cite{zaldarriaga_seljak_1998} and therefore their detection -- a source of constraints on the matter perturbation evolution at redshift $z\sim1$ when light massive neutrinos and elusive dark energy both play potentially visible roles. Very recently, the {\it direct} detection of the lensing-induced $B$-mode and the primordial $B$-mode has been reported by the {\sc polarbear} experiment \cite{polarbearpaper} and the {\sc bicep2} experiment \cite{biceppaper}, respectively. In the standard cosmological paradigm, the $TB$ end $EB$ cross-correlations are vanishing. However, they remain important quantities to be estimated from the data. This is because on the one hand, these odd-parity cross-spectra are comprehensive, end-to-end, null tests of the presence of instrumental and/or astrophysical systematic effects still present in the data (see {\it e.g.} \cite{hu_etal_2003,yadav_etal_2010}). On the other hand, as some non-standard cosmological mechanisms could produce nonvanishing odd-parity cross-spectra, their detection could become a smoking gun of such effects with potentially far-reaching consequences for our understanding of the Universe. Examples of such mechanisms include a primordial stochastic magnetic field, which generates $TB$ and $EB$ correlations, if this magnetic field possesses a helical component \cite{pogosian_etal_2002,caprini_etal_2004,kahniashvili_etal_2005} or a pseudoscalar inflaton field which naturally couples to the electromagnetic field in a parity-dependant way \cite{sorbo_2011,anber_sorbo,cook_sorbo}. Similar effects can be obtained due to a rotation of the plane of linear polarization of the CMB photons traveling from the last scattering surface to our detectors. This could result from either the Faraday rotation induced by interaction with background magnetic fields \cite{kosowsky_loeb_1996,kosowsky_etal_2005,campanelli_etal_2004,scoccola_etal_2004} or interactions with pseudoscalar fields on the trajectory of CMB photons \cite{carroll_1998}. In this paper, we consider the case of odd-parity angular power spectra as probes of parity violation in the primordial Universe as induced by gravity. The implication of chiral gravity on CMB anisotropies has been first explored in Ref. \cite{lue_etal_1999} and then in Ref. \cite{contaldi_etal_2008} where it was shown that if parity is violated by gravitation at the linear level, CMB polarized anisotropies should exhibit non vanishing $EB$ and $TB$ cross-correlations. This idea has been theoretically strengthened in Refs. \cite{magueijo_benincasa_2011,bethke_magueijo_2011a,bethke_magueijo_2011b}, and the idea that gravity could be parity dependent can be traced back to its formulation by {\it e.g.} Cartan and Kibble \cite{kibble_1961} or Ashtekar \cite{ashtekar_1986}. The possible detection of such parity asymmetry using CMB datas coming from a satellite-like mission has been discussed in Refs. \cite{xia_2012,saito_2007}, in Ref. \cite{wang_etal_2013} in the Horava-Lifshitz framework and in Ref. \cite{gluscevic_etal_2010} (including the case of a ballon-borne experiment in the latter). We amend and elaborate on this proposal of Refs.\cite{lue_etal_1999,contaldi_etal_2008,xia_2012,saito_2007,gluscevic_etal_2010} in three directions. First, chiral gravity leads to {\it primary} $TB$ and $EB$ cross-correlations which are latter on, deformed by the weak gravitational lensing by large scale structure. As this could potentially lead {\it e.g.} $EE$ correlations to leak into $EB$ correlations (which would partially mask the primary $EB$), we therefore include in the predicted $C_\ell$'s the impact of lensing. Second, we make use of a Fisher matrix formalism to assess the potential detection of chiral gravity from the measurements of CMB polarized anisotropies in two typical experimental setups: small-scale experiments as motivated by operating (or forthcoming) balloon-borne or ground-based experiments such as \textsc{polarbear}, \textsc{sptpol}, {\sc qubic} or \textsc{actpol}, for ground-based experiments \cite{ground}, and, such as {\sc spider} or \textsc{ebex}, for balloon-borne experiments \cite{balloon}, and, satellite-like missions as motivated by {\it e.g.} {\sc l}ite\textsc{bird}, \textsc{prism} or \textsc{pix}i\textsc{e} proposals \cite{satellite}. Estimation of the uncertainties on the reconstructed $C^{TB(EB)}_\ell$ (subsequently used in the Fisher matrix) is based first on a na\"\i{ve} mode-counting (as a reference), and, second, on Monte-Carlo simulations coupled to a {\it realistic} statistical, pure pseudospectrum based estimators of angular power spectra. Thirdly, we assess the impact of a miscalibration of the orientation of the polarized detectors which creates spurious $TB$ and $EB$ correlations coming from $TE$ and $EE,~BB$ respectively. The paper is organized as follows. The section \ref{sec:cell} is devoted to the theoretical prediction of the $TB$ and $EB$ angular power spectra including the impact of weak gravitational lensing by large scale structure. We present the statistical uncertainties on the reconstruction of $C^{TB(EB)}_\ell$'s using pure pseudospectrum estimators in Sec. \ref{sec:uncer}. The results of the application of such an approach to the two above-defined typical cases of CMB experiments dedicated to polarization, small-scale experiments and satellite-like missions, are presented in Secs. \ref{sec:satres} and \ref{sec:smares} respectively. We finally conclude and discuss the potential detection of chiral gravity within CMB anisotropies in the last section, Sec. \ref{sec:concl}, and discuss the relevance and extension of those results to other possible sources of parity violation in the primordial universe. The technical details are provided in the appendices \ref{app:lensing} and \ref{appb}. | \label{sec:concl} In this paper, we investigate the constraints that could be set on chiral gravity from the detection of the CMB $TB$ and $EB$ correlations, taking into account statistical uncertainties as incurred by pure pseudospectrum reconstruction of the CMB angular power spectra and considering the impact of miscalibrating the orientation of the polarized detectors. (We stress that all the constraints we have set are for positive valued $r_{(-)}$. They however equally apply to negative values of $r_{(-)}$ as in practice, the derived constraints are for $\left|r_{(-)}\right|$.) We have shown that such a detection of parity violation leading to non zero $C^{TB(EB)}_\ell$ is beyond the scope of forthcoming small-scale measurements of CMB polarized anisotropies. Even in the most optimistic case of 100\% of parity violation and a tensor-to-scalar ratio of 0.2, and underestimating the uncertainties by using a mode-counting approach, the signal-to-noise ratio on the amplitude of parity asymmetric tensor mode is only of $\sim1.2$, and it rapidly diminishes to values smaller than unity for smaller values of the tensor-to-scalar ratio, $r_{(+)}$, or a smaller percentage of parity violation. This is because most of the constraints come from the largest angular scales which cannot be measured with enough significance by those experiments. Moreover, even in the case of vanishing $TB$ and $EB$ cross-correlations, the statistical uncertainties on their reconstruction via pure pseudospectrum estimators lead to an upper bound of the level of parity violation of more than 100\% at 95\% CL. Since this level is theoretically bounded from above at 100\%, this means that no significant constraint can be set on this type of parity violation using datas from ongoing or forthcoming small-scale experiments. In the case of a potential satellite mission dedicated to primordial $B$-mode, we have shown that a detection with at least $2\sigma$ is possible for 100\% of parity violation and a tensor-to-scalar ratio of at least 0.05. A 1$\sigma$ detection is still achieved for 50\% of parity violation and a tensor-to-scalar ratio of at least 0.05 and a $2.5\sigma$ detection would be possible for $r_{(+)}=0.2$. We found that by a measurement of vanishing $TB$ and $EB$ angular power spectra using pure pseudospectrum estimators, the level of parity violation is bounded from above: $\left|\delta\right|\leq0.92$ at 95\% CL. We have also shown that for such an experimental configuration where sampling variance is dominating at the largest scales -- precisely those scales which allows for constraining parity violation --, the impact of self-calibrating the miscalibration angle could have a significant impact on the final estimation of the level of parity violation, the reported signal-to-noise ratio being degraded by a factor of $\sim1.09$ for a miscalibration angle of 0.1 degree and, more significantly, by a factor $\sim2.4$ for an angle of 1 degree. In this very last case, a $2\sigma$ detection of parity violation becomes possible only for $\delta=1$ and $r_{(+)}=0.2$. We stress that such an impact is revealed in the context of the pseudospectrum estimation of the angular power spectra. By making use of the na\"\i{v}e mode-counting expression for the covariance of the reconstructed $C_\ell$'s, it is formally shown that self-calibration of the orientation of the polarizers does not impact on the significance of the reconstruction of $r_{(-)}$ assuming that the covariance is dominated by sampling variance and that one have access to the entire set of cross-correlations between the six estimated angular power spectra. Nevertheless, this last assumption is broken by pseudospectrum estimators (because of mode-mixing and binning) leading to degeneracies between $r_{(-)}$ and $\Delta\psi$. In the context of chiral gravity from the Ashtekar formulation of general relativity, the parameter $\delta$ amounting the level of parity breaking is related to the imaginary Barbero-Immirzi parameter via $\delta=2i\gamma/(1-\gamma^2)$ \cite{bethke_magueijo_2011a}. (We will restrict to the case of purely imaginary values of $\gamma$ though the formalism can be extended to the case of any arbitrary complex values of $\gamma$ \cite{bethke_magueijo_2011b}.) The (absolute) level of parity breaking is encoded in $\left|\delta\right|$ leading to $\left|\gamma\right|=\left(1\pm\sqrt{1-\delta^2}\right)/\left|\delta\right|$. Considering the statistical error bars from a pseudospectrum reconstruction of the $C_\ell$'s, detecting $\left|\gamma\right|=1$ is possible using datas from a satellite mission with a statistical significance ranging from $2.3\sigma$ to $5.4\sigma$ for a tensor-to-scalar ratio ranging from $0.05$ to $0.2$, respectively. Assuming a detection of $\left|\delta\right|=0.5$ translates into a detectable value of $\gamma=0.26$ or $\gamma=3.73$, meaning that such a form of chiral gravity is detectable with CMB polarized anisotropies if $0.26\leq\left|\gamma\right|\leq3.75$. The significance of that detection for a future satellite mission ranges from $1.1\sigma$ to $2.5\sigma$ for a tensor-to-scalar ratio of 0.05 and 0.2. Detecting $TB$ and $EB$ angular power spectra which are consistant with zero leads to an upper bound on the parity violation level $\left|\delta\right|\leq0.92$ at 95\% CL for $r_{(+)}=0.05$ and $\delta\leq0.39$ at 95\% CL for $r_{(+)}=0.2$. This would mean that $0.66\leq\left|\gamma\right|\leq1.5$ is excluded at 95\% CL for $r_{(+)}=0.05$ (the exclusion range at 68\% CL would be $0.24\leq\left|\gamma\right|\leq4.1$), and that $0.2\leq\left|\gamma\right|\leq4.9$ is excluded at 95\% CL for $r_{(+)}=0.2$ (the exclusion range at 68\% CL would be $0.098\leq\left|\gamma\right|\leq10.1$). In the context of a pseudoscalar inflaton, the amount of parity violation is given by \cite{cook_sorbo} \begin{equation} \left|\delta\right|=\frac{8.6\times10^{-7}\left(\frac{H^2}{2M^2_\mathrm{Pl}}\frac{e^{4\pi\xi}}{\xi^6}\right)}{1+8.6\times10^{-7}\left(\frac{H^2}{2M^2_\mathrm{Pl}}\frac{e^{4\pi\xi}}{\xi^6}\right)} \nonumber \end{equation} and \begin{equation} r_{(+)}=8.1\times10^{7}\left(\frac{H^2}{M^2_\mathrm{Pl}}\right)\left[1+8.6\times10^{-7}\left(\frac{H^2}{2M^2_\mathrm{Pl}}\frac{e^{4\pi\xi}}{\xi^6}\right)\right]. \nonumber \end{equation} For a given value of $r_{(+)}$, one can express $\frac{H^2}{2M^2_\mathrm{Pl}}$ as a function of $\frac{e^{4\pi\xi}}{\xi^6}$ and plug it into $\left|\delta\right|$. Following Ref. \cite{cook_sorbo}, one introduces the parameter $\tilde{X}=e^{2\pi\xi}/\xi^3$ which is related to $r_{(+)}$ and $\delta$ via \begin{equation} \tilde{X}=\left(\frac{1.37\times10^7}{\sqrt{r_{(+)}}}\right)\sqrt{\left(\frac{\left|\delta\right|}{1-\left|\delta\right|}\right)\left(1+\frac{\left|\delta\right|}{1-\left|\delta\right|}\right)}. \nonumber \end{equation} As compared to Ref. \cite{cook_sorbo}, $\tilde{X}$ is related to their $X$ parameter via $X=\epsilon\times\tilde{X}$ with $\epsilon$ the first slow-roll parameter. For $r_{(+)}=0.05$, one obtains the following upper bound on $\tilde{X}$: $\tilde{X}\leq73\times10^7$ at 95\% CL. For $r_{(+)}=0.2$, this upper bound is strengthened to $\tilde{X}\leq3\times10^{7}$ at 95\% CL. This has to be compared to the upper bound reported in Ref. \cite{cook_sorbo}: $\tilde{X}\leq6\times10^7$ at 95\% CL, using the upper bound set by Planck on primordial non-gaussianities, $f_\mathrm{NL}<150$ \cite{planck_gauss}. This means that constraining such models with $TB$ and $EB$ is on par with the constraints that can be set with measurements of non-gaussianities assuming a rather high value of $r_{(+)}$. | 14 | 4 | 1404.6660 |
1404 | 1404.2333_arXiv.txt | { The onset of the asymmetry in planetary nebulae (PNe) occurs during the short transition between the end of the asymptotic giant branch (AGB) phase and the beginning of the PN phase. Sources in this transition phase are compact and emit intensely in infrared wavelengths, making high spatial resolution observations in the infrared mandatory to investigate the shaping process of PNe. Interferometric VLTI IR observations have revealed compelling evidence of disks at the cores of PNe, but the limited sensitivity, strong observational constraints, and limited spatial coverage place severe limits on the universal use of this technique. Inspired by the successful detection of proto-planetary disks using spectro-astrometric observations, we apply here for the first time this technique to search for sub-arcsecond structures in PNe. Our exploratory study using CRIRES (CRyogenic high-resolution Infra-Red EchelleSpectrograph) commissioning data of the proto-PN IRAS\,17516$-$2525 and the young PN SwSt\,1 has revealed small-sized structures after the spectro-astrometric analysis of the two sources. In IRAS\,17516$-$2525, the spectro-astrometric signal has a size of only 12\,mas, as detected in the Br$\gamma$ line, whereas the structures found in SwSt\,1 have sizes of 230\,mas in the [Fe~{\sc iii}] line and 130\,mas in the Br$\gamma$ line. The spectroscopic observations required to perform spectro-astrometry of sources in the transition towards the PN phase are less time consuming and much more sensitive than VLTI IR observations. The results presented here open a new window in the search of the small-sized collimating agents that shape the complex morphologies of extremely axisymmetric PNe. } | Using long-slit high-resolution spectra, the SA technique enables us to resolve small-sized structures in different astronomical objects. Inspired by these results, we have investigated the use of this technique in the search of structures at mas scales at the innermost regions of sources in the transition from the AGB to the PN phase. For this purpose, we have used CRIRES commissioning data of the proto-PN IRAS\,17516$-$2525 and the young PN SwSt\,1 to develop the methodology and the tools to perform such SA analyses. Our SA analysis has been able to detect the presence of a small-sized structure, $\sim$12 mas in size, in IRAS\,17516$-$2525 that can be interpreted as small bipolar lobes. As for SwSt\,1, the SA analysis shows two structures with different sizes, $\sim$130-230 mas, and inclinations in the [Fe\,{\sc iii}] and Br$\gamma$ lines. Tailored SA observations at different PAs with its corresponding calibrations, as well as a proper model fitting, are mandatory to derive a conclusive interpretation of the SA signatures. Remarkably, the SA analysis based on CRIRES commissioning data has been able to detect structures which are comparable in size to those resolved by MIDI-VLTI in Mz\,3 and M\,2-9, and more compact than those detected in AFGL\,915 with radio-interferometric observations at Plateau de Bure. Compared to MIDI-VLTI, CRIRES SA observations are more sensitive, not restricted by fixed baselines, and the data interpretation is less model dependent. We conclude that the SA technique is very suitable for the search of small-sized disks and asymmetric structures in the short transition from the AGB to the PN phase. | 14 | 4 | 1404.2333 |
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1404 | 1404.2934_arXiv.txt | Recent work has exploited pulsar survey data to identify temporally isolated, millisecond-duration radio bursts with large dispersion measures (DMs). These bursts have been interpreted as arising from a population of extragalactic sources, in which case they would provide unprecedented opportunities for probing the intergalactic medium; they may also be linked to new source classes. Until now, however, all so-called fast radio bursts (FRBs) have been detected with the Parkes radio telescope and its 13-beam receiver, casting some concern about the astrophysical nature of these signals. Here we present FRB~121102, the first FRB discovery from a geographic location other than Parkes. FRB~121102 was found in the Galactic anti-center region in the 1.4-GHz Pulsar ALFA survey with the Arecibo Observatory with a DM = $557.4\; \pm\; 3\; \DMunits$, pulse width of $3\; \pm 0.5$~ms, and no evidence of interstellar scattering. The observed delay of the signal arrival time with frequency agrees precisely with the expectation of dispersion through an ionized medium. Despite its low Galactic latitude ($b = -0.2^{\circ}$), the burst has three times the maximum Galactic DM expected along this particular line-of-sight, suggesting an extragalactic origin. A peculiar aspect of the signal is an inverted spectrum; we interpret this as a consequence of being detected in a sidelobe of the ALFA receiver. FRB~121102's brightness, duration, and the inferred event rate are all consistent with the properties of the previously detected Parkes bursts. | \label{intro} Radio pulsar surveys sample the sky at high time resolution and are thus sensitive to a range of time variability and source classes. Over the last decade, there has been renewed interest in expanding the purview of pulsar search pipelines, which traditionally exploit the periodic nature of pulsars, to also search for single dispersed pulses. This led to the discovery of Rotating Radio Transients \citep[RRATs;][]{mll+06}, which are believed to be pulsars that are either highly intermittent in their radio emission or have broad pulse-energy distributions that make them more easy to discover using this technique \citep{wsrw06}. Of the now nearly 100 known RRATs\footnote{\url{http://astro.phys.wvu.edu/rratalog/}}, the vast majority emit multiple detectable pulses per hour of on-sky time, though a few have thus far produced only one observed pulse \citep{bb10,bbj+11}. The dispersion measures (DMs) of the RRATs are all consistent with a Galactic origin, according to the NE2001 model for Galactic electron density \citep{cl02}. Single-pulse search methods have also discovered a new class of fast radio bursts (FRBs) in wide-field pulsar surveys using the 13-beam, 1.4-GHz receiver at the Parkes radio telescope \citep{lbm+07, kskl12, tsb+13}. Most have been found far from the Galactic plane and have DMs that are anomalously high for those lines-of-sight. \citet{lbm+07} reported the first such burst with a Galactic latitude of $b = -42^{\circ}$ and DM = $375$\,$\DMunits$. The expected DM contribution for that line-of-sight from the ionized interstellar medium (ISM) in our Galaxy is only $25$\,$\DMunits$ according to the NE2001 model. The DM excess has been interpreted as coming from the ionized intergalactic medium (IGM) and led to the conclusion that FRBs are extragalactic. More recently, \citet{tsb+13} reported four FRBs with Galactic latitudes of $|b| > 40^{\circ}$ and DMs ranging from 521~to~$1072\, \DMunits$. The expected Galactic DM contribution along the lines-of-sight of these bursts is $30 - 46$\,$\DMunits$, i.e.\ only $3-6$\,\% of the observed DM can be attributed to our Galaxy. An additional FRB candidate was reported by \citet{kskl12} and is at a lower Galactic latitude ($b = -4^{\circ}$) than the other five reported Parkes FRBs. This source could be of Galactic origin given that the measured DM = $746$\,$\DMunits$ is only 1.3 times the maximum expected DM from NE2001 along this line-of-sight. The dispersion delay of all of the published FRBs are consistent with the expected $\nu^{-2}$ dispersion law. Additionally, the burst reported by \citet{lbm+07} showed frequency-dependent pulse broadening that scaled as $\nu^{-4.8 \pm 0.4}$, consistent with the expected value of $-$4.0 to $-$4.4 \citep{lr99} for scattering by the ISM. The brightest burst reported by \citet{tsb+13} showed a clear exponential tail and a pulse duration that scaled as $\nu^{-4.0 \pm 0.4}$. This provides additional credence to the interpretation that the signal is of astrophysical origin. Generally, FRBs have been found in minute to hour-long individual observations; multi-hour follow-up observations at the same sky positions have thus far failed to find repeated bursts. Thus, FRBs are considered a different observational phenomenon from RRATs based on DMs in excess of the predicted Galactic contribution and the fact that none of the FRBs has been seen to repeat. {\bfref At this point, however, we can not be certain that the bursts are non-repeating. Detecting an astrophysical counterpart will be an important step in determining whether we expect repeated events.} The progenitors and physical nature of the FRBs are currently unknown. The FRBs have brightness temperatures well in excess of thermal emission ($T_b >10^{33}$ K) and therefore require a coherent emission process. One possible source of repeating, extragalactic FRBs is extremely bright, rare Crab-like giant pulses from extragalactic pulsars, which repeat over much longer time scales than currently constrained. Proposed extragalactic sources of non-repeating, fast radio transients include evaporating primordial black holes \citep{r77}, merging neutron stars \citep{hl01}, collapsing supramassive neutron stars \citep{fr13}, and superconducting cosmic strings \citep{cssv12}. Alternatively, \citet{lsm13} suggest a repeating, Galactic source - flares from nearby, magnetically active stars, in which the DM excess is due to the star's corona. In this scenario, additional pulses could also be observed and potentially at a different DM. Localizing FRBs with arcsecond accuracy is technically challenging but will help identify potential host galaxies, or stars, and multi-wavelength counterparts. In any pulsar survey, the vast majority of statistically significant signals are due to man-made radio frequency interference (RFI), which can originate far from the telescope or be locally generated. RFI can also mimic some of the characteristics of short-duration astronomical signals. Thus, care is needed when interpreting whether a particular signal is astronomical in origin, and claims of a new source class require due consideration and skepticism. {\bfref The situation is further complicated by the discovery of ``perytons" \citep{bbe+11}. Perytons are which are short duration radio bursts observed in pulsar surveys over a narrow range of DMs but have patchy spectra and are observed in many beams simultaneously. } The fact that FRBs have so far been observed with only the Parkes telescope has raised some concern - even though the observed brightness distribution and event rate can explain why other experiments have so far not detected any similar signals. In this article we report the discovery of an FRB with the Arecibo Observatory. The FRB was found as part of the Pulsar ALFA (PALFA) survey of the Galactic plane \citep{cfl+06}. {\bfref This detection, made with a different telescope at a different geographic location, bolsters the astrophysical interpretation of a phenomenon seen until now only with Parkes.} The outline for the rest of this paper is as follows. In \S \ref{survey}, we describe the PALFA survey and the observations that led to the discovery of the new FRB. The burst's properties are discussed in \S \ref{burst} and the implied FRB event rate is described in \S \ref{rate}. A discussion of the possible origin of this FRB, both astrophysical and otherwise, is outlined in \S \ref{origin}. In \S \ref{discussion} we discuss the implications of our discovery for FRBs in general and present our conclusions. | \label{discussion} Under the assumption that FRB~121102 is extragalactic in origin, and following \citet{tsb+13}, we can estimate the redshift, $z$, of the burst based on the observed total dispersion delay across the bandpass, which is fortuitously for ALFA $\sim$1~ms for each 1~$\DMunits$ of DM, i.e.\ $\Delta t_{\rm obs}=552$~ms (Figure~\ref{fig:frb}). The contributions to $\Delta t_{\rm obs}$ are from: (i) free electrons in the Galactic ISM ($\Delta t_{\rm ISM}$); (ii) the intergalactic medium ($\Delta t_{\rm IGM}$); (iii) the putative host galaxy ($\Delta t_{\rm Host}$). Adopting the NE2001 electron density for this line-of-sight, we find $\Delta t_{\rm ISM}\simeq 184$~ms. To be consistent with the estimates presented in \citet{tsb+13}, we make the same assumptions about $\Delta t_{\rm IGM}$ and $\Delta t_{\rm Host}$ as presented in their Figure~S3. Using the DM scaling relationship for the intergalactic medium model of \citet{i04}, we find $\Delta t_{\rm IGM}\simeq 1200\, z$~ms. For a host galaxy DM contribution of 100~$\DMunits$, $\Delta t_{\rm Host} \simeq 100$~ms$/(1+z)^2$. The condition $\Delta t_{\rm obs} = \Delta t_{\rm ISM} + \Delta t_{\rm IGM} + \Delta t_{\rm Host}$ is met when $z=0.26$. The redshift value can be taken as an upper bound because it is plausible that a host galaxy can contribute more to the total delay than we have assumed. As was the case for the redshift estimates presented by \citet{tsb+13} ($z$=0.45 to 0.96), the contributions to $\Delta t_{\rm obs}$ are highly model dependent, and therefore the $z$ value should be used with caution. With these caveats in mind, the implied co-moving radial distance at $z=0.26$ would be $D \sim 1$~Gpc. The FRB pseudo-luminosity $S D^2 \sim 1 \times 10^{12}$~Jy~kpc$^2$ and energy output is $\sim 10^{38}$~ergs for isotropic emission and $\sim 10^{37}$~ergs for emission beamed over 1~steradian. Both values are consistent with the FRBs from \citet{tsb+13}. Using this estimate of the co-moving distance, we can also constrain the brightness temperature, i.e.\ $T_b \sim 1 \times 10^{34}\; D_{\rm Gpc}^2\; \rm K$, where $D_{\rm Gpc}$ is the source distance in units of Gpc. This unphysically large brightness temperature requires a coherent emission process. All of the FRBs observed to date show less temporal scattering than pulsars with similar DMs. Using a population of pulsars at low Galactic latitudes, \citet{bcc+04} determined an empirical relation for pulse broadening timescale versus DM and observing frequency. For example, the predicted pulse broadening timescales for $\rm DM=500$~to~$1000\; \DMunits$ are 2~to~2000\,ms at 1.4\,GHz, albeit with a large scatter in the observed distribution. By comparison, only FRB~110220 has a measurable scattering timescale of $\sim 5$ ms with $\rm DM = 910\; \DMunits$ \citep{tsb+13}, roughly a factor of 200 less than predicted by \citet{bcc+04}. Using this single FRB scattering measurement, \citet{lkmj13} scale the \citet{bcc+04} relation. The scaled relation predicts a scattering timescale for FRB~121102 of $\sim$0.04 ms, which is shorter than the time resolution of the data. If this relation can in fact be applied broadly to FRBs, then it is not surprising that we detected no scattering. \citet{lkmj13} point out that for a given scattering screen, the largest observed scattering occurs when the screen is near the mid-point between the source and observer due to geometric effects. This suggests that the IGM or an intervening galaxy located midway along the line-of-sight would be the most important contribution to the scattering of FRBs. However, \citet{cm03} show that for a source imbedded in a region of high scattering, for example near the center of the host galaxy or for a line-of-sight that passes through the host's galactic disk, the observed scattering can still be dominated by the host galaxy even at large distances. Observations of scattering along extragalactic lines-of-sight by \citet{lof+08} and more theoretical calculations by \citet{mk13} suggest that scattering in the IGM is several orders-of-magnitude lower than in the ISM, which is consistent with the observations of FRBs. One caveat to the conclusions of this paper is that the search presented here, and all searches for dispersed radio bursts, are optimized for signals with a $\nu^{-2}$ dispersive time delay. This simple approach introduces a selection effect in what signals breach the S/N threshold used to identify candidates, and thus which signals are deemed worthy of close inspection. We note, however, that this selection effect is not severe in our case as the fractional bandwidth we have used here (20\%) is just barely sufficient to see the quadratic curvature of the burst delay. Although the poor localization of FRB~121102 prevents a detailed search for a multi-wavelength counterpart, we searched for any major high-energy events that were both contemporaneous and co-located on the sky. We checked the Gamma-ray Coordinates Network (GCN) archive of $\gamma$-ray bursts and found no potential association with FRB~121102. There are no plausible associations with X-ray transients detected by current all-sky monitors, and there are no observations of the field of FRB~121102 (within a 20$^{\prime}$ radius) with X-ray telescopes. There is no source associated with this object in either the \emph{ROSAT} All Sky Survey Catalog or the \emph{Fermi} Source Catalog. In summary, we have described the Arecibo discovery of FRB~121102, a single, highly dispersed pulse in the PALFA survey. This is the first claimed FRB detection that has been found with a telescope other than Parkes. The large DM excess, roughly three times what would be expected from the Galactic ISM along this line-of-sight, the absence of repeat bursts, and the low interstellar scattering suggest that this is an FRB and not a Galactic emitter such as an RRAT. Using the occurrence rate inferred from the PALFA discovery, we predict that, in the coming years, PALFA will find two to three more FRBs in the remaining outer Galaxy survey region. | 14 | 4 | 1404.2934 |
1404 | 1404.7499_arXiv.txt | Determining the energy spectrum of an electron population can give key insights into the underlying physics of a radio source; however, the lack of high resolution, broad-bandwidth observations has left many ambiguities in our understanding of radio galaxies. The improved capabilities of telescopes such as the JVLA and LOFAR mean that within the bandwidth of any given observation, a detailed spectral shape can now be produced. We present recent investigations of powerful FR-II radio galaxies at GHz and MHz frequencies and show for the first time their small-scale spectral structure. We highlight problems in traditional methods of analysis and demonstrate how these issues can now be addressed. We present the latest results from low frequency studies which suggest a potential increase in the total energy content of radio galaxy lobes with possible implications for the energetics of the population as a whole. | \label{introduction} In principle, a region emitting synchrotron radiation will preferentially cool higher energy electrons leading to a steeper, more strongly curved spectrum in older regions of plasma. Models of this `spectral ageing' have become a commonly used tool when describing the processes involved in the lobes of FR-I and FR-II type radio galaxies. Historically, investigations of spectral ageing in radio sources have largely been limited by the fact that only a few narrow-band frequencies have been available. However, the new generation of radio telescopes has changed this situation dramatically. The increased sensitivity, bandwidth and dynamic range provided by instruments such as LOFAR and the JVLA now allow for spectral curvature to be determined across an entire frequency band of a single pointing. This type of detailed spectral analysis is set to become standard practice, hence it is vital that methods are developed to handle new broadband radio observations. The `Broadband Radio Analysis ToolS' (BRATS) was therefore developed to provide a range of analysis, statistical and model fitting tools for broadband radio data \citep{harwood13}. Within these proceedings, we detail how these new methods have been applied to observations at GHz frequencies, and describe our current research at low-frequencies as part of the LOFAR nearby AGN key science project (KSP). | 14 | 4 | 1404.7499 |
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1404 | 1404.6236.txt | {We re-examine large scalar fields within effective field theory, in particular focussing on the issues raised by their use in inflationary models (as suggested by BICEP2 to obtain primordial tensor modes). We argue that when the large-field and low-energy regimes coincide the scalar dynamics is most effectively described in terms of an asymptotic large-field expansion whose form can be dictated by approximate symmetries, which also help control the size of quantum corrections. We discuss several possible symmetries that can achieve this, including pseudo-Goldstone inflatons characterized by a coset $G/H$ (based on abelian and non-abelian, compact and non-compact symmetries), as well as symmetries that are intrinsically higher dimensional. Besides the usual trigonometric potentials of Natural Inflation we also find in this way simple {\em large-field} power laws (like $V \propto \phi^2$) and exponential potentials, $V(\phi) = \sum_{k} V_k \; e^{-k \phi/M}$. Both of these can describe the data well and give slow-roll inflation for large fields without the need for a precise balancing of terms in the potential. The exponential potentials achieve large $r$ through the limit $|\eta| \ll \epsilon$ and so predict $r \simeq \frac83(1-n_s)$; consequently $n_s \simeq 0.96$ gives $r \simeq 0.11$ but not much larger (and so could be ruled out as measurements on $r$ and $n_s$ improve). We examine the naturalness issues for these models and give simple examples where symmetries protect these forms, using both pseudo-Goldstone inflatons (with non-abelian non-compact shift symmetries following familiar techniques from chiral perturbation theory) and extra-dimensional models.} \begin{document} | The very early universe seems to have been a remarkably simple place: all we know --- and we now know a fair amount --- about the properties of primordial fluctuations is consistent with the predictions of the simplest single-field inflationary models \cite{Inflation}. Although there are an impressive number of single-field inflationary models \cite{Single-field-Zoo}, an even more impressively large number of them lay bleeding on the ground after the recent discovery of primordial gravitational waves \cite{BICEP2}. Although BICEP2 finds the tensor-to-scalar ratio is $r = 0.2^{+ 0.07}_{- 0.05}$, it is likely that better modeling of foregrounds will reduce this value (for instance the preliminary analysis for one such model gives $r = 0.16^{+0.06}_{-0.05}$ \cite{BICEP2}). To be conservative, for the purposes of this paper we simply take the BICEP2 observations to imply % \be \label{BICEP2r} r_{\rm exp} \gsim 0.1 \,. \ee What makes this interesting is that a great many models do not give $r$ this large, and so would be decisively ruled out if the observation is confirmed. Among other things this includes the majority of (but not all) string-inflation models \cite{StrInfPostPlanck}. One way to see why large $r$ is such a challenge is the Lyth bound \cite{LythBound} that relates a value this large for $r$ to a trans-Planckian range through which the inflaton rolls in simple single-field models; something which many of the known models do not do. Although efforts have been made towards evading the Lyth bound \cite{LythUnbound}, we believe that it should be embraced: if large field displacements are difficult (but not impossible) to obtain then the fact that Nature seems to want them is likely to be very informative. Within simple inflationary models primordial perturbations ultimately arise as quantum fluctuations in the gravitational field. Are large field excursions consistent with the validity of the semiclassical approximations on which controlled quantum gravity calculations rely? In principle they need not be inconsistent: in the end the semiclassical approximation relies on the low-energy approximation \cite{EFTGrav}, but it is not necessary that large field values must cost a large energy density. Flat directions in supersymmetric field theories --- for which large fields cost precisely zero energy --- provide perhaps the most direct existence proof that large fields and large energies need not be linked. What is required for a controlled calculation is an understanding of the behaviour of the lagrangian in the large-field regime, including the behaviour of the scalar potential and the inflaton's target-space geometry. If the scalar potential is bounded for large fields, what is its asymptotic behaviour? And what is the large-field limit of the target-space metric? Are these asymptotic forms stable under quantum corrections? Symmetries and approximate symmetries can help with all of these issues. For instance, if the inflaton is a pseudo-Goldstone boson for an approximate symmetry \cite{pGB}, then in the symmetry limit it becomes a bona-fide Goldstone boson on which the scalar potential cannot depend. Furthermore, for a symmetry breaking pattern where $G$ breaks to a subgroup $H$ the Goldstone fields parameterize the coset space $G/H$ \cite{CCWZ}, on which the $G$-invariant target-space metrics can also be systematically identified, allowing a coherent picture of the full target-space geometry in both the large- and small-field regimes. These features are not changed appreciably if $G$ is only an approximate symmetry, since by assumption the geometry of the target space is only slightly perturbed. Furthermore, this approximate symmetry can also ensure the whole picture survives quantum effects. But since the scalar potential itself is a symmetry-breaking effect, any regime where it grows without bound (such as for large fields) introduces a worry about the validity of the entire approximate-symmetry picture. Natural inflation \cite{NatInf} --- with a trigonometric potential naturally arising from the weak breaking of an abelian shift symmetry: $\phi \to \phi + \omega \, f$ --- provides the simplest example along these lines (for which the potential is everywhere bounded), but does not exhaust the possibilities. In this paper we describe a broader class of potentials that are similarly protected by (generalized) shift symmetries, and so for which the inflaton is a pseudo-Goldstone boson. In particular we display simple examples that illustrate the following two more general points: % \begin{itemize} \item They show that pseudo-Goldstone bosons can enjoy more complicated potentials than the simple trigonometric potential of Natural Inflation. In particular we show that exponential potentials generically arise (all the while keeping a positive definite target-space metric) when non-compact symmetries are considered. (Ours is not an exhaustive study, and more general kinds of potentials than trigonometric or exponential are also likely possible.) % \item Because the geometry of one dimension is not that interesting, the constraints imposed by symmetry on the large-field target-space geometry are most informative when there is more than one pseudo-Goldstone boson. In this case symmetry arguments can dictate the target-space geometry, showing how field redefinitions can be used to relate the large- and small-field regimes (trading complications in the potential for complications in the scalar kinetic terms). We work through the illustrative example of two Goldstone bosons for which the target space is the coset $SU(1,1)/U(1)$. Of course PLANCK and BICEP2 are also informative here since they also constrain the existence of other light fields during inflation through their contributions to isocurvature perturbations \cite{QCDAxion}. % \end{itemize} In general a proper understanding of the target space geometry and the nature of the small symmetry breaking terms allows an understanding of what kinds of asymptotic forms are appropriate to the scalar potential in the large-field regime.\footnote{The usual formulation of large-field inflation often is cast using specific coordinates in field space, obscuring the freedom to perform field redefinitions that map large fields to small. Appendix \ref{app:fieldredef} briefly reviews how to recast inflationary slow-roll conditions in a more covariant way.} In particular it asks whether or not the potential diverges in the large-field limit (as is assumed when parameterizing it as a positive power of $\phi$), and if it does how quickly it does. For instance, in the $SU(1,1)/U(1)$ example we find that the potential might diverge or be bounded, but in either case naturally admits an expansion in powers of exponentials, $e^{-k \phi/M_p}$, for large $\phi$. The rest of the paper is set out as follows. Next, in \S\ref{sec:phenomenology}, we review the inflationary phenomenology of single-field models built on the simplest forms for large-field potentials: those that look like $V \propto \phi^p$; $V \sim V_0 - V_1/\phi^q$ and $V \sim V_0 - V_1 \, e^{-k\phi}$. We show how all of these can successfully describe the combined observations (including $r \gsim 0.1$) and identify which parts of parameter space can do so. Then \S\ref{sec:naturalness} identifies the naturalness issues these models face, and show in particular two different ways that symmetries --- 4D generalized shift symmetries, very similar to Natural Inflation, as well as extra-dimensional symmetries --- can generate exponential potentials in the large-field limit. \S\ref{sec:naturalness} also briefly describes some of the proposed UV completions for these models and comments on the parameter ranges that are found in explicit examples. Finally, our conclusions are summarized in \S\ref{sec:conclusions}. We collect useful material in three appendices. In Appendix A we present a covariant formulation of slow-roll conditions including an invariant definition of large field for multi-field models. Appendix B presents explicit discussion of simple compact and non-compact cosets $SO(3)/SO(2)$ and $SO(2,1)/SO(2)$ including the construction of invariant metrics and explicit scalar potentials (exponentials and power-law) following standard techniques of chiral perturbation theory. Appendix C is a discussion of the supersymmetric $SU(1,1)/U(1)$ coset including a discussion of $D$ and $F$ terms. | \label{sec:conclusions} This seems to be the decade of fundamental scalar fields in particle physics. If confirmed, the results of BICEP2 open a golden window onto the high energies relevant to UV completions of the standard models of particle physics and cosmology. The fact that these results point to high energies, of order the GUT scale, hints at new physics close to the Planck scale. Having the simplest slow-roll inflation successfully capture CMB observations while also generating sufficiently large tensor modes puts strong constraints on specific inflationary models. In particular, the indication that trans-Planckian values of the scalar fields may have been explored is likely to be quite informative. Although these need not be problematic for a controlled EFT approach (provided the large-field energy densities are not too large), there is little in our experience at small fields that is guaranteed to go over as well to the properties of physics at trans-Planckian field values. We argue here a point of view (that we believe is widely appreciated) that a good way to formulate large-field physics in a precise and controlled way is if the theory is close to a symmetric limit that controls the large-field asymptotic forms. Such a symmetry can also protect against large quantum corrections and thereby help control the validity of the EFT of interest. Turning this argument around, the BICEP2 results may be hinting at the existence of approximate symmetries of whatever UV complete theory applies at the high energies at which we now know inflation must take place. This makes it important to explore systematically the possible symmetries that can play a role in this regard, and in this article we explore several kinds that can be relevant: spontaneously broken approximate global symmetries (for which the inflaton is a pseudo-Goldstone boson), supersymmetry and extra-dimensional spacetime symmetries. For pseudo-Goldstone inflatons, we are led to new kinds of inflaton potentials in addition to the standard shift/axionic symmetries that have long been studied. By studying both non-abelian and non-compact coset spaces we identify several kinds of natural potentials, including both a class of exponential potentials. These exponential potentials can describe observations well, including allowing $r \gsim 0.1$. They make it difficult to obtain $r$ too much larger (such as being as large as 0.2) when $n_s \simeq 0.96$, and so could easily become ruled out as the errors on $r$ and $n_s$ improve. We also show how the simplest quadratic potential of chaotic inflation can be included in the class of natural pseudo-inflaton models. This becomes possible because pseudo-Goldstone inflation allows the potential to be Taylor expanded for small fields whenever $\phi \ll F$ (generically leading to quadratic behaviour near a minimum), but this small-field regime can include trans-Planckian fields when the decay constant satisfies $F > M_p$ (as it typically does in inflationary models). Finally, we indicate how these natural potentials (including in particular exponential inflation) often emerge from UV completions, and in particular are generic in supergravity and string theory. This helps to have a broader perspective in the search for concrete UV complete models of inflation. Finding a fully UV satisfactory realization of all features of inflation, including modulus stabilization remains an open question. | 14 | 4 | 1404.6236 |
1404 | 1404.0121.txt | %{}{Incomplete}{Incomplete}{Incomplete}{} %context {In the framework of the Herschel/PRISMAS Guaranteed Time Key Program, the line of sight to the distant ultracompact \HII\ region W51e2 has been observed using several selected molecular species. Most of the detected absorption features are not associated with the background high-mass star-forming region and probe the diffuse matter along the line of sight. We present here the detection of an additional narrow absorption feature at $\sim$ 70 km~s$^{-1}$ in the observed spectra of HDO, NH$_3$ and \cthree. The 70\,\kms\ feature is not uniquely identifiable with the dynamic components (the main cloud and the large-scale foreground filament) so-far identified toward this region. The narrow absorption feature is similar to the one found toward low-mass protostars, which is characteristic of the presence of a cold external envelope. The far-infrared spectroscopic data were combined with existing ground-based observations of $^{12}$CO, $^{13}$CO, CCH, CN, and C$_3$H$_2$ to characterize the 70\,\kms\ component. Using a non-LTE analysis of multiple transitions of NH$_3$ and CN, we estimated the density ($n(H_2)\sim$(1--5)\,10$^5$\,\cmcub) and temperature (10--30\,K) for this narrow feature. We used a gas-grain warm-up based chemical model with physical parameters derived from the NH$_3$ data to explain the observed abundances of the different chemical species. We propose that the 70\,\kms\ narrow feature arises in a dense and cold clump that probably is undergoing collapse to form a low-mass protostar, formed on the trailing side of the high-velocity filament, which is thought to be interacting with the W51 main cloud. While the fortuitous coincidence of the dense clump along the line of sight with the continuum-bright W51e2 compact HII region has contributed to its non-detection in the continuum images, this same attribute makes it an appropriate source for absorption studies and in particular for ice studies of star-forming regions.} | The high-sensitivity large-scale CO maps of giant molecular clouds provided the first evidence for the filamentary nature of the interstellar medium \citep{ungerechts1987,bally1987}. The far-infrared all-sky IRAS survey and several mid-infrared surveys (ISO, Spitzer/MIPS) revealed the ubiquitous filamentary structure of both the dense and diffuse ISM \citep{low1984}. Most recently, the unprecedented sensitivity and large-scale mapping capabilities of Herschel/SPIRE have revealed a large network of parsec-scale filaments in Galactic molecular clouds and have suggested an intimate connection between the filamentary structure of the ISM and the formation of dense cloud cores \citep{andre2010,molinari2010}. Here we present evidence of the detection of a dense clump possibly formed by the interaction of an extended foreground filament with the giant molecular cloud W51. Located at a distance of 5.41$^{+0.31}_{-0.28}$\,kpc \citep{sato2010}, W51 is a radio source with a complicated morphology in which many compact sources are superposed on extended diffuse emission \citep{bieging1975}. The line of sight to W51 intersects the Sagittarius spiral arm nearly tangentially (l = 49\arcdeg), which means that sources over a $\sim 5$\,kpc range of distances are superimposed on the line of sight. Based on the $^{12}$CO and $^{13}$CO 1--0 line emission of these sources \citet{carpenter1998} divided the molecular gas associated with the W51 HII region into two subgroups: a giant molecular cloud (1.2$\times 10^6$ M$_{\odot}$) at $\sim$ 61~\kms, and an elongated (22$\times$136\,pc) molecular cloud akin to a filamentary structure (1.9$\times 10^5$ M$_{\odot}$) at 68~\kms. While the brightest radio source at 6\,cm (G49.5-0.4) is spatially and kinematically coincident with the W51 giant molecular cloud, the G49.2-0.3, G49.1-0.4, G49.0-0.3, and G48.9-0.3 radio sources seem to be associated with the 68~\kms\ cloud. \citet{carpenter1998} speculated that the massive star formation activity in this region resulted from a collision between the W51 giant molecular cloud and the high-velocity (68~\kms) cloud. G49.5-0.4 is the brightest source in the W51 main region. The continuum emission from the ultracompact \HII\ region was collectively named W51e before it was resolved into compact components, labeled as W51e1 to W51e8 \citep{scott1978,gaume1993,mehringer1994,zhang1997}. Figure~\ref{fig_overview} shows the 70\,\micron\ continuum emission as observed with PACS by Hi-Gal in color, along with contours of $^{13}$CO (1--0) emission integrated between 68.9 and 70.3\,\kms\ to locate the GMC and the filament\footnote{The PACS data shown in Fig.~\ref{fig_overview} are a Level 2 data product produced by the standard data reduction pipeline for PACS-SPIRE parallel mode observations as obtained from the Herschel Science Archive. The $^{13}$CO data have taken from the Galactic Ring Survey done by FCRAO available online at http://www.bu.edu/galacticring/}. The continuum emission appears to be better correlated with the filament, possibly suggesting star formation activity in the filament. \begin{figure} \begin{center} %\includegraphics[width=0.47\textwidth]{../w51overview.eps} \includegraphics[width=0.47\textwidth]{Figs/w51filament.eps} \caption{Color image shows 70\,\micron\ continuum emission observed with PACS. The white contours represent $^{13}$CO (1--0) emission integrated between 68.9 and 70.3\,\kms\ corresponding to the ``filament", observed using FCRAO as part of the Galactic Ring Survey. The black cross shows the position of the HIFI observations. \label{fig_overview}} \end{center} \end{figure} As part of the {\em Herschel} key program ``PRobing InterStellar Molecules with Absorption line Studies (PRISMAS) we have observed several selected spectral lines in absorption toward the source W51e2 to study the foreground material along the line of sight using HIFI. These observations have for the first time detected in absorption a component at 70\,\kms\ (a velocity even higher than the velocity of the 68\,\kms\ filament) along with the 57\,\kms\ source velocity component, in \cthree, HDO, and NH$_3$. The observation of \cthree\ was motivated by the importance of small carbon chains for the chemistry of stellar and interstellar environments either as building blocks for the formation of long carbon chain molecules, or as are products of photo-fragmentation of larger species such as PAHs \citep{radi1988,pety2005}. \cthree\ has been identified with {\em Herschel}/HIFI for the first time in the warm envelopes of star-forming hot cores like W31C, W49N and DR21(OH) \citep{mookerjea2010,mookerjea2012}. Within the PRISMAS program, the ground-state HDO transition at 894\,GHz has only been detected at the velocity of the HII regions. Based on observations of high-mass star-forming regions \citep{jacq1990,pardo2001, vandertak2006, persson2007, bergin2010} and low-mass protostars \citep{fcliu2011, coutens2012,coutens2013}, the HDO/H$_2$O ratio remains lower than the D/H ratio derived for other deuterated molecules observed in the same sources although it is clearly higher than the cosmic D/H abundance. This low deuterium enrichment of water could provide an important constraint to the astrochemical models to explain the chemical processes involved in the formation of water. Ammonia (NH$_3$) is a key species in the nitrogen chemistry and a valuable diagnostic because its complex energy level structure covers a very broad range of critical densities and temperatures. It has been observed throughout the interstellar medium ever since its detection in 1968 \citep{cheung1968} but mainly in its inversion lines at cm wavelengths. With \emph{Herschel} it became possible to observe ground-state rotational transitions with high critical densities at THz frequencies of both ortho and para symmetries, for instance, in cold envelopes of protostars \citep{hilyblant2010}, in diffuse/translucent interstellar gas \citep{persson2012}, and hot cores \citep{neill2013}. We have combined the HIFI observations with IRAM observations of CCH, \cthreehtwo\ and CN in all of which the 70\,\kms\ feature is detected in absorption. We have used Local Thermodynamic Equilibrium (LTE) and non-LTE models to derive the physical parameters of the dense clump at 70\,\kms\ and its relation to the W51 molecular cloud. We also present a gas-grain, warm-up chemical model that consistently explains the abundances of all the species observed in the clump. This paper is organized as follows: Sections 2 and 3 describe the Herschel and IRAM observations, respectively. Section\,4 describes the complex velocity structure observed along the sightline to W51e2. Section\,5 discusses the column and volume densities derived from all the tracers that detect the 70\,\kms\ clump. In Sec. 6 we compare the observed abundances of the various chemical species with a grain warm-up based chemical model. Section\,7 discusses various derived properties of the 70\,\kms\ clump and presents a possible formation scenario for the clump. | In the direction of W51e2 we have for the first time detected HDO and \cthree\ in a dense molecular clump at 70\,\kms\ that is not directly associated with hot star-forming cores. \cthree\ in dense clouds has so far only been detected in the warm envelopes associated with star-forming cores. Abundance of \cthree\ in dense clouds is explained via warming-up of grain mantles to release CH$_4$ and subsequently C$_3$H$_2$, the first step in the chemical pathway for the formation of \cthree, as explained in detail in {\citet{hassel2011}} and {\citet{mookerjea2012}}. Thus, detection of \cthree\ in the dense clump at 70\,\kms\ otherwise not identified as a star-forming region is as surprising as the detection of deuterated water in the cloud that is physically not related to the W51 region. HDO is most likely formed in the ice mantles within cold ($<$20\,K) translucent clouds \citep{cazaux2011} in regions shielded from UV radiation. Indeed, at $A_{\rm v} \gtrsim 3$ \citep{whittet1988}, species from the gas phase accrete onto dust grains and initiate ice formation. After the formation of water and deuterated water in the ice mantles in the star formation scenario, the environment undergoes gravitational collapse and the local density increases, leading to the formation of a protostar. The surrounding medium then heats up and releases the icy mantles into the gas phase, leading to a significant HDO abundance in the gas phase \citep{coutens2012,coutens2013b}. Based on the non-LTE modeling of the five observed NH$_3$ transitions, we characterize the region of the that filament gives rise to the 70\,\kms\ feature as having $T_{\rm kin}$ = 30\,K and $n({\rm H_2})$ = 5$\times 10^5$\,\cmcub. The abundances predicted by the grain warm-up chemical model reproduce the observed abundances to within factors of 2 to 3 at all times later than 3.5 Myr for all species except for \cthreehtwo. %For HDO the agreement between %the observed and model abundances is for a somewhat shorter duration %at $t> 3.5$\,Myr. \subsection{CCH/\cthreehtwo\ ratio in the clump} Previous observations of CCH and \cthreehtwo\ toward dense dark clouds and Photon Dominated Regions (PDRs) derived a tight correlation between the two species. The ratio $N$(CCH)/$N$(\cthreehtwo) is $\sim 61$ in OMC-1 \citep{blake1987} and it ranges between 10 and 25 in PDRs such as Horsehead nebula and IC63 whereas in dark clouds such as TMC-1, and L134N the ratio is $\sim 1$ \citep{teyssier2004}. For the dense clump detected toward W51e2 the observed $N$(CCH)/$N$(\cthreehtwo) of 67 is closer to the value observed in the dense active star-forming region OMC-1 than to the values seen in UV-illuminated clouds and dark clouds. \subsection{Deuterium enrichment in the clump} Using Herschel/HIFI, a number of transitions of H$_2$O and H$_2$$^{18}$O species corresponding to a range of excitations have been observed for the first time. The 70\,\kms\ feature is detected unambiguously in the H$_2$O 2$_{0,2}$--1$_{1,1}$ transition as well as the H$_2$$^{18}$O 1$_{1,1}$--0$_{0,0}$ transition \citep{flagey2013}. However, an accurate determination of the total H$_2$O column density is compromised by the crowded environment along the line of sight with high optical depth, as well as by partial blending of the components. The column density for the less abundant para-H$_2$$^{18}$O species is estimated to be (7 $\pm$ 1) 10$^{11}$ cm$^{-2}$ for the 70\,\kms. The abundance ratio between $^{16}$O and $^{18}$O is commonly accepted to be $\sim$ 560 \citep{wilson1994}, although significant variation has been observed throughout the Galaxy. Accordingly, we estimate a D/H ratio in water vapor of about 9.6 $\times$ 10$^{-4}$ (assuming an ortho-para ratio of 3), similar to the value found in Orion\,KL by \citet{neill2013}. This ratio has also been estimated from HDO observations of the low-mass protostar IRAS 16293-2422 \citep{coutens2012,coutens2013}. The derived values of HDO/H$_2$O were 1.8$\times 10^{-2}$ in the hot core, 5$\times 10^{-3}$ in the cold envelope, and about 4.8$\times 10^{-2}$ in the external photodesorption layer. The value estimated from our analysis is consistent with the value found in the cold envelope of the IRAS 16293-2422 low-mass star-forming region, in which the ices are not affected by thermal desorption or photo-desorption by the FUV field. For the best-fit grain warm-up model that reproduces the observed abundances of the other molecules, at times later than 3.5 Myr the HDO/H$_2$O ratio drops from 0.002 to 0.001 for an initial HD/H$_2$ ratio of $1.5\times 10^{-5}$. The calculated HDO/H$_2$O ratio is consistent with the observed ratio of 9.6 $\times$ 10$^{-4}$. Although HDO forms in the medium density, pre-collapse phase, it mostly remains on grain mantles, since the nonthermal desorption mechanisms have a limited efficiency. In these models, which reach only 30\,K, the release of HDO into the gas phase is attributed to nonthermal processes related to cosmic-ray interactions rather than warming up or direct FUV photodesorption. Although the chemical model predicts an HDO/H$_2$O ratio consistent with the observed value for $t>3.5$\,Myr, it overestimates the HDO/H$_2$O ratio by more than a factor of 10 for earlier times $t<1$\,Myr. The calculated HDO/H$_2$O ratio is substantially enriched above the initial HD/H$_2$ ratio, primarily because of the grain surface formation route of HDO. The ratio is furthermore enriched above the observed value because of the lower column density of H$_2$O calculated by the chemical models. We explored the possibility of enhanced cosmic-ray ionization rates. For enhanced cosmic ray rates the chemical models produce abundances very different from the observed values for all molecules. This result for the 70\,\kms\ cloud is in contrast to the higher values of $\zeta$ deduced by \citet{indriolo2012} along the same line of sight for the absorption features arising because of diffuse foreground. This implies that the absorbing gas is not diffuse, but is dense with a standard value of $\zeta$. \subsection{Possible origin of the clump} The 70\,\kms\ absorption dip caused by the foreground material in the direction of W51e2, although not associated with a star-forming core, shows chemical abundances typically found in the envelopes of low-mass protostellar candidates. We propose that the feature arises in a dense ($n$(H$_2$)=(1-5)$\times 10^5$\,\cmcub) and cold (10-30\,K) clump that is formed within the much larger scale filament (detected in CO) deemed to be interacting with the W51 main molecular cloud. A possible scenario for the formation of the dense clump can be outlined based on the collision of the filament with W51 main \citep{kang2010}. In this scenario, cloud-cloud collision lead to the compression of the interface region and initiate the formation of stars as seen in W51 \citep{habe1992}. The molecular clumps at the interface are heated, but the molecular clumps on the trailing side remain cold and hence appear in absorption. Models for collision between two different clouds suggest that such collisions disrupt the larger cloud while the small cloud is compressed and subsequently forms stars. It is possible that the dense clump detected in absorption is formed on the trailing side of the filament that interacts with the main molecular cloud. Based on its density and somewhat elevated temperatures ($\sim 30$\,K), it is also possible that this cloud is also collapsing to eventually form stars. In the absence of any observational evidence, we currently do not consider the source to be an IRDC. We suggest that the nondetection of this star-forming clump in the continuum is due to its chance coincidence along the line of sight with the much stronger W51e2 continuum source. Higher spectral and angular resolution observations of the foreground gas responsible for the absorption dip at $\sim 70$\,\kms\ are needed to understand its spatial distribution. We also propose to explore a clearer detection of the protostar using mid-infrared observations of the strong spectral features of molecular ices toward this region. | 14 | 4 | 1404.0121 |
1404 | 1404.4874_arXiv.txt | The dense interiors of massive galaxies are among the most intriguing environments in the Universe. In this paper we ask when these dense cores were formed and determine how galaxies gradually assembled around them. We select galaxies that have a stellar mass $>3\times 10^{10}$\,\msun\ inside $r=1$\,kpc out to $z=2.5$, using the 3D-HST survey and data at low redshift. Remarkably, the number density of galaxies with dense cores appears to have decreased from $z=2.5$ to the present. This decrease is probably mostly due to stellar mass loss and the resulting adiabatic expansion, with some contribution from merging. We infer that dense cores were mostly formed at $z>2.5$, consistent with their largely quiescent stellar populations. While the cores appear to form early, the galaxies in which they reside show strong evolution: their total masses increase by a factor of $2-3$ from $z=2.5$ to $z=0$ and their effective radii increase by a factor of $5-6$. As a result, the contribution of dense cores to the total mass of the galaxies in which they reside decreases from $\sim 50$\,\% at $z=2.5$ to $\sim 15$\,\% at $z=0$. Because of their early formation, the contribution of dense cores to the total stellar mass budget of the Universe is a strong function of redshift. The stars in cores with $M_{\rm 1\,kpc}>3\times 10^{10}$\,\msun\ make up $\sim 0.1$\,\% of the stellar mass density of the Universe today but 10\,\% -- 20\,\% at $z\sim 2$, depending on their IMF. The formation of these cores required the conversion of $\sim 10^{11}$\,\msun\ of gas into stars within $\sim 1\,$kpc, while preventing significant star formation at larger radii. | The central regions of massive elliptical galaxies such as NGC\,1399 and NGC\,4472 are different from any environment seen in galaxies such as the Milky Way. The mean stellar densities are $\sim 10$\,\msun\,pc$^{-3}$ in the central kpc, and their velocity dispersions reach or even exceed $\sim 300$\,\kms. The stellar populations are old, metal rich, and strongly $\alpha$-enhanced, indicating that the stars were formed early in a short, intense period of star formation ({Franx} \& {Illingworth} 1990; {Worthey}, {Faber}, \& {Gonzalez} 1992; {Davies}, {Sadler}, \& {Peletier} 1993; {Kuntschner} {et~al.} 2001, 2010, and many other studies). Star formation in these central regions likely took place under very different physical conditions than those in the present-day disk of the Milky Way, possibly leading to a bottom-heavy IMF with an excess of low mass stars compared to the Milky Way IMF ({van Dokkum} \& {Conroy} 2010; {Treu} {et~al.} 2010; {Krumholz} 2011; {Cappellari} {et~al.} 2012; {Conroy} \& {van Dokkum} 2012; {Hopkins} 2013). These dense centers also host the most massive black holes in the Universe ({Magorrian} {et~al.} 1998; {Ferrarese} \& {Merritt} 2000; {Gebhardt} {et~al.} 2000), which probably accreted most of their mass during the peak star formation epoch. Despite their high star formation efficiency in the past, dense regions are hostile to star formation today: quiescence correlates well with velocity dispersion and with stellar surface density ({Kauffmann} {et~al.} 2003; {Franx} {et~al.} 2008; {Wake}, {van Dokkum}, \& {Franx} 2012; {Bell} {et~al.} 2012). The dense interiors of massive galaxies account for only a small fraction of the total stellar mass in the present-day Universe, but given the old ages of their stars this fraction is expected to increase with redshift. In fact, the formation of the dense central parts of elliptical galaxies may preceed the assembly of the rest of the galaxies. Many quiescent galaxies at $z=1.5 - 2.5$ are much more compact than nearby galaxies of the same mass (e.g., {Daddi} {et~al.} 2005; {Trujillo} {et~al.} 2006; {van Dokkum} {et~al.} 2008; {Cimatti} {et~al.} 2008; {Damjanov} {et~al.} 2009; {Williams} {et~al.} 2010), and as first shown by {Bezanson} {et~al.} (2009) the central densities of the compact high redshift galaxies are broadly similar to those of massive elliptical galaxies today. This is consistent with the idea that massive galaxies have grown inside-out since $z\sim 2$, with their cores forming at higher redshift and their outer envelopes building up slowly through star formation, minor mergers, or other processes (e.g., {Loeb} \& {Peebles} 2003; {Bezanson} {et~al.} 2009; {Naab}, {Johansson}, \& {Ostriker} 2009; {van Dokkum} {et~al.} 2010; {Hopkins} {et~al.} 2010; {Oser} {et~al.} 2010; Feldmann et al.\ 2010; {Szomoru} {et~al.} 2013). In this paper we focus exclusively on these dense central regions of massive galaxies: we ask what their number density is, what their contribution is to the overall stellar mass density, and how the galaxies that they are part of were built up around them. In practice, we select galaxies out to $z=2.5$ that have $\log M_{\rm 1\,kpc}\gtrsim 10.5$, that is, a stellar mass exceeding $3.2 \times 10^{10}$\,\msun\ within a sphere of radius $r=1$\,kpc.\footnote{We refer to the region within this radius as the ``core'' throughout this paper, realizing that this may cause confusion. The same term has been used extensively in the literature to describe the surface density profile of early-type galaxies on much smaller scales (e.g., {Faber} {et~al.} 1997).} We do not limit the sample to quiescent galaxies but select all objects that satisfy this stellar density criterion. Our approach is different from studies of the properties of galaxies at fixed total stellar mass, or fixed number density. In fact, as we show in \S\,\ref{ndens.sec} the evolution of the number density of galaxies with $\log M_{\rm 1\,kpc}>10.5$ is different from that of the general population of massive galaxies. Our study is more closely related to the work of {Bezanson} {et~al.} (2011) on the evolution of the velocity dispersion function; Bezanson et al.\ converted observed effective radii, stellar masses, and {Sersic} (1968) indices to velocity dispersions whereas we convert the same parameters to a stellar mass within a physical radius of 1\,kpc. In this paper we do not make any a priori selection on star formation rate or galaxy size. Nevertheless, this paper has implications for the evolution of massive quiescent galaxies at $z\sim 2$. It is generally thought that these galaxies have grown substantially in size over the past 10 Gyr, but this interpretation is complicated by the fact that the number density of quiescent galaxies has increased by an order of magnitude over this time period ({Brammer} {et~al.} 2011; {Cassata} {et~al.} 2013; {Muzzin} {et~al.} 2013a). As discussed by {van Dokkum} {et~al.} (2008), {van der Wel} {et~al.} (2009), Trujillo et al.\ (2011), {Newman} {et~al.} (2012), {Carollo} {et~al.} (2013), {Szomoru} {et~al.} (2013), and others, the evolution of the mass-size relation of quiescent galaxies could be partially driven by the continuous addition of large, recently quenched star-forming galaxies, in which case the growth of individual quiescent galaxies would be smaller than that of the population. Some studies have even suggested that compact quiescent galaxies barely evolve at all (e.g., {Poggianti} {et~al.} 2013). As we show in \S\,\ref{ndens.sec} the evolution of galaxies with dense cores appears to require substantial evolution in the sizes and masses of individual compact galaxies after $z\sim 2$. The paper is structured as follows. In \S\,\ref{data.sec} we describe the sources of data that are used. In \S\,\ref{select.sec} the selection of galaxies with dense cores is described. Sections \ref{ndens.sec} and \ref{buildup.sec} form the heart of the paper. In \S\,\ref{ndens.sec} the ``core mass function'' is discussed, that is, the number of galaxies as a function of their mass within 1 kpc. This section also presents the evolution of the cumulative number density of galaxies with $\log M_{\rm 1\,kpc}>10.5$, and interprets the evolution in the context of various physical processes. Finally, it places the total stellar mass locked up in dense cores in the context of the evolving stellar mass density of the Universe. In \S\,\ref{buildup.sec} the properties of galaxies that have dense cores are analyzed; here we show that the core-hosting galaxies likely evolved significantly since $z\sim 2$, increasing both their total mass and (particularly) their effective radii. We also discuss the nature of star forming galaxies with dense cores. The paper is summarized in \S\,\ref{conclusions.sec}. | \label{conclusions.sec} In this paper we have identified dense cores in galaxies out to $z=2.5$, using data from the 3D-HST project augmented by low redshift information from UltraVISTA and the Sloan Digital Sky Survey. We find that the evolution of cores with mass $\log(M_{\rm 1\,kpc}) \sim 10.5$ is well described by mild mass loss, suggesting that their stars form a passive stellar population since $z\sim 2.5$. We note that mergers may also contribute to the evolution, and that the effects of mass loss are sensitive to the assumption that 100\,\% of the stellar ejecta mix with the hot halo gas. At $z\sim 2.5$ the cores make up $\sim 50$\,\% of the total mass of the galaxies that they are part of. At lower redshift they make up a decreasing fraction of the total mass, and by $z=0$ they are embedded in large envelopes of stars with effective radii $\sim 5$\,kpc. We focused on cores of a fixed high mass, but we note that the evolution of the core mass function is mass-dependent (see Fig.\ \ref{densfunc.fig}), with low mass cores showing strong positive evolution in their number density. This mass dependence has also been seen in the total mass function ({Marchesini} {et~al.} 2009) and in the velocity dispersion function ({Bezanson} {et~al.} 2011). At low masses star formation may lead to a relatively uniform build-up of galaxies, with the stellar density increasing at all radii, whereas at high masses galaxies are built up inside-out (see {van Dokkum} {et~al.} 2013). The negative mass evolution of the cores has consequences for the interpretation of massive star forming galaxies at $z=1-2.5$ and the evolution of quiescent galaxies, as discussed in \S\,\ref{buildup.sec}. However, we emphasize that not all massive galaxies have dense cores: selecting on total mass produces different samples than selecting on core mass, as is obvious in Figs.\ \ref{massmass.fig} and \ref{massfrac.fig}. Our conclusions only hold for galaxies with a dense core, and leave open the possibility that massive galaxies with low core masses have different evolutionary trajectories. It so happens that by $z\sim 2$ our selection mostly overlaps with the population of massive, quiescent galaxies at that redshift, which is why we can rule out several proposed models for their evolution (see \S\,\ref{buildup.sec}). We also find that, at fixed total mass and redshift, the presence of a dense core is a good predictor of quiescence and (perhaps more interestingly) its absence is a nearly perfect predictor of star formation (see Fig.\ \ref{massfrac.fig}). The latter result is strikingly unambiguous: of 91 galaxies with $M_{\rm tot}>10^{11}$\,\msun\ and $M_{\rm 1\,kpc}<10^{9.5}$\,\msun\ only one is quiescent. Apparently the presence of a dense core is a ``non-negotiable'' requirement for stopping star formation in massive galaxies. Perhaps the most important result of this paper is that the contribution of stars in dense cores to the stellar mass density of the Universe increases strongly with redshift, reaching values of 10\,\% -- 20\,\% at $z\sim 2$ (\S\,\ref{contrib.sec} and Fig.\ \ref{fractions.fig}b). In light of this high fraction we suggest that the formation of these cores is an important aspect of star formation, galaxy formation, and black hole formation at high redshift. Interestingly it is not yet clear how this happened. Near the end of their main star formation epoch, prior to stellar mass loss, the cores were even more massive and compact than at $z\sim 2$. The gas mass that was converted to stars inside 1\,kpc must have approached $10^{11}$\,\msun. Furthermore, this gas must have arrived in the core without forming many stars at larger radii: the quiescent descendants at $z\sim 2$ have small effective radii and no low surface brightness envelopes (e.g., {Szomoru} {et~al.} 2010, 2013). Several mechanisms have been proposed for creating very compact massive galaxies, such as mergers ({Hopkins} {et~al.} 2008) and disk instabilities ({Dekel} \& {Burkert} 2014). However, reproducing the surface density profiles of the cores has proven to be challenging (see {Wuyts} {et~al.} 2010). It will also be interesting to see whether models can be created that simultaneously explain the existence of large, massive disks such as that of M101 and of extremely compact cores of similar mass. Forming large disks requires feedback and significant angular momentum (e.g., {Guedes} {et~al.} 2011), whereas forming dense cores requires rapid cooling and a mechanism to lose angular momentum efficiently ({Sales} {et~al.} 2012; {Dekel} \& {Burkert} 2014). Whatever the mechanism is for getting gas into the center, the core mass will build up quickly when star formation begins. The adiabatic enhancement discussed in Appendix \ref{lossaper.sec} should also apply ``in reverse'': when mass is added to the center, the mass within 1\,kpc will increase as $\sim (M'/M)^2$ due to adiabatic contraction. Whether forming dense cores have been observed is a matter of debate. As discussed in \S\,\ref{buildup.sec} star forming galaxies with dense cores, such as those identified by {Patel} {et~al.} (2013) and {Barro} {et~al.} (2013), may not be forming the core itself but stars away from the center. Spatially-resolved star formation maps (e.g., {F{\"o}rster Schreiber} {et~al.} 2011; {Nelson} {et~al.} 2013; {Wuyts} {et~al.} 2013), or spectroscopy to determine the kinematics of the gas, may provide more information on the location of star formation in these objects. Given the high metallicity of the centers of present-day elliptical galaxies and the high densities, the star forming cores must have had very large amounts of absorption. They may be largely invisible in the optical and near-IR, and possibly even at larger wavelengths.\footnote{Somewhat akin to dragonflies, which are aquatic during their nymph stage.} Studies of red, far-IR selected galaxies have shed some light on this issue (e.g., {Tacconi} {et~al.} 2006, 2008; {Wang}, {Barger}, \& {Cowie} 2012; {Gilli} {et~al.} 2014). The ``prototype'' would be a dusty star forming galaxy with a compact morphology and a gas dispersion that matches the dispersion of present-day elliptical galaxies; such an object has recently been identified (Nelson et al.\ 2014). The main uncertainty in the analysis is the conversion of light to mass. As discussed in Appendix A and elsewhere, the systematic uncertainties are $\sim 0.1$\,dex, or half of the observed evolution in the core mass. Stellar kinematics are a crucial check on the mass measurements (see, e.g., {Bezanson} {et~al.} 2013; {van de Sande} {et~al.} 2013; {Belli} {et~al.} 2013), although models for the structure of the galaxies and their dark matter are required to interpret them. Furthermore, we have ignored radial gradients in $M/L$ ratio. Our analysis shows that the core masses do not grow but the total masses do, which means the stellar populations in the core are likely different from those at larger radii. The available evidence suggests that these gradients are generally small ({Szomoru} {et~al.} 2013), but it is difficult to measure them at the relevant spatial scales: 1 kpc corresponds to a single native WFC3 pixel. Spatially-resolved studies of strongly-lensed galaxies with dense cores could address this issue. | 14 | 4 | 1404.4874 |
1404 | 1404.3665_arXiv.txt | The reconstruction of a (non)canonical scalar field Lagrangian from the dark energy Equation of State (EoS) parameter is studied, where it is shown that any EoS parametrization can be well reconstructed in terms of scalar fields. Several examples of EoS parameters are studied and the particular scalar field Lagrangian is reconstructed. Then, we propose some new parametrizations that may present a (fast) transition to a phantom dark energy EoS (where $w_{DE}<-1$) and the scalar field Lagrangian is also reconstructed numerically. Furthermore, the properties of these parametrizations of the dark energy EoS are studied by using supernovae Ia data (HST Cluster Supernova Survey) combined with Standard Ruler datasets [Cosmic Microwave Background (CMB) and Baryon Acoustic Oscillations (BAO)] and its comparison with the $\Lambda$CDM model is analyzed. Then, the best fit of the models is obtained, which provides some information about whether a phantom transition may be supported by the observations. In this regard, the crossing of the phantom barrier is allowed statistically but the occurrence of a future singularity seems unlikely. | In 1998 a deviation on the luminosity distance of Supernovae Ia (Sne Ia) was observed by two independent groups \cite{SN1}, a fact that was interpreted as the acceleration of the universe expansion. Later on, other independent observations such as the Cosmic Microwave Background (CMB) \cite{WMAP}-\cite{Ade:2013zuv} or the Baryon Acoustic Oscillations (BAO) \cite{Eisenstein:2005su} have confirmed such hypothesis, which has been widely accepted by the scientific community since then. Then, a large number of candidates, enclosed under the name of dark energy, have been proposed in order to understand the mechanism that produces such accelerating expansion (for a review on dark energy candidates see \cite{R-DE}). The list of models includes a cosmological constant, canonical/phantom scalar fields \cite{Quintessence}, vector fields \cite{VT} or modifications of General Relativity (GR) \cite{modified-gravity}, among others. \\ Moreover, an interesting and useful approach for analyzing dark energy models is aimed to study the dark energy equation of state (EoS) as an effective description instead of reconstructing theoretical models. In this sense, dynamical EoS's that deviate from the cosmological constant have been widely studied, where perfect fluids with inhomogeneous EoS and redshift-dependent parameters have been proposed which may accomplish the late-time acceleration, and even the entire cosmological history by unifying the dark energy epoch and the inflationary phase (see Ref.~\cite{Nojiri:2005sr}). Moreover, an effective description of the behavior of the dark energy EoS simplifies the fit of the free parameters while comparing with observational data, such that theoretical models, as modified gravities or scalar-tensor theories, can be tested by using effective parametrizations of the EoS along the period of interest of the universe evolution. In this regard, several parametrizations of the dark energy EoS have been proposed over the last decade and its comparison with observational data has been studied (see Refs.~\cite{Huterer:2000mj}-\cite{Lazkoz:2007cc}). Some of these models lead to $\Lambda$CDM as the one with major statistical support but also other possibilities are allowed. Furthermore, the possibility that dark energy behaves as a phantom fluid, whose effective EoS parameter would turn out $w_{DE}<-1$, has been also widely explored in the literature (see Ref.~\cite{phantom}) in spite of that such transition may lead to large instabilities in some particular phantom models \cite{vikman}. Such kind of EoS produces a phase of super-accelerating expansion that may end in a future singularity (for a classification of future singularities, see Refs.~\cite{Nojiri:2005sx}-\cite{barrow1}), whose analysis has attracted much interest, as may content important information on the structure of spacetime and its topology (see Ref.~\cite{barrow}). Hence, singular cosmologies have been explored within several frameworks, including modified gravities (see Ref.~\cite{Nojiri:2008fk}). Furthermore, observations seem not to discard every phantom scenario and even some analysis highly support such possibility when studying carefully the observational data \cite{Lazkoz:2006gp}.\\ In the present paper, we present a reconstruction method for the action of a (non)canonical scalar field by just specifying the EoS parameter. Then, some examples of parametrizations of the dark energy EoS are reconstructed in terms of the scalar field. Such reconstruction method may be extended to other theoretical models as modified gravities to obtain the gravitational action from the dark energy EoS. The aim of such reconstruction is to provide a method to relate a phenomenological description, as the dark energy EoS, with the underlying theory that leads to such phenomenological behavior. Then, we propose some new EoS parametrizations that experience fast changes, and which may give rise to fast crossings of the phantom barrier and eventually to the occurrence of a future singularity. The best fit of the free parameters of the models are found by using some observational datasets from Standard rulers (CMB \cite{Spergel:2006hy} and BAO \cite{Eisenstein:2005su}) and Sne Ia \cite{Suzuki:2011hu}. The comparison of the results obtained by using each dataset is analyzed as well as the comparison with the $\Lambda$CDM model. The value of the relative matter density $\Omega_m^0$ is found to be very close to the $\Lambda$CDM model, while the best fit of the EoS parameters does not discard the transition to the phantom epoch but disfavor the occurrence of future singularities. \\ The paper is organized as follows: section \ref{reconstruction} deals with the reconstruction of scalar field models from the dark energy EoS. Then, section \ref{NewEoS} is devoted to the analysis of some new parametrizations of the EoS, where a preliminary study of the cosmological evolution and the occurrence of future singularities are analyzed. Section \ref{ObsData} deals with the fit of the free parameters of the model with observational data and its comparison with $\Lambda$CDM. Finally, in section \ref{Conclusions}, we discuss the results of the paper. | \label{Conclusions} In this paper, we have focused on the analysis of some parametrizations of the dark energy EoS, where we have implemented a new method to reconstruct a scalar field Lagrangian that gives rise to a particular EoS parameter. Parametrizing the dark energy EoS is commonly used to describe effectively the underlying theoretical model, since it facilitates the analysis of the dark energy EoS and its confrontation with the observational data. Within this aim we have focused on the reconstruction of a simple theoretical model, a (non)canonical scalar field, starting from the EoS parameter, where several examples have been studied. Then, two new parametrizations have been proposed, analyzed in terms of a scalar field and compared with the observational data. The aim of the proposed parametrizations has been to study the possibility of a fast transition and in particular the possibility of crossing the phantom barrier. Note that both models contain $\Lambda$CDM as a special case, when $w_0=0$, in which case the scalar field action reduces to a cosmological constant term, as shown in section \ref{reconstruction}\\ Then, by using Supernova Ia and Standard Ruler data, both models have been analyzed and also compared with $\Lambda$CDM. The results show that both parametrizations leads to a $\chi^2_{min}$ value that is in general slightly smaller than the $\Lambda$CDM model whereas the value of the matter density yields $\Omega_m^0=0.25$ at the best fit. Besides, the resulting $\chi^2_{red}$ value within both models is below respect the resulting one for the $\Lambda$CDM model when using both standard candles as standard ruler data. In addition, the second parametrization (\ref{2.2}) leads to better results regarding the value of the $\chi^2_{min\ (red)}$ in comparison with the parametrization (\ref{2.1}).\\ Nevertheless, both parametrizations contain more free parameters than the $\Lambda$CDM model and specifically the parameter $z_0$ presents a large indetermination, specially when dealing with $z_0$ values that can not be constrained with experimental data. Furthermore, by analyzing the different approaches computed along the paper, the contour plots of both models show that $w_0\sim 0$ is very likely, in spite of the best fit is slightly displaced from $\Lambda$CDM. In addition, it is remarkable that both models do not lead to future singularities at the best fit, as shown in Table \ref{table5}, and neither in most of the previous results, although the possibility of the occurrence of a future singularity is not excluded within the confidence region of the contour plots, as depicted in Figs.~\ref{fig:SN.1}-\ref{fig:C.2}. \\ Indeed, while analyzing the first model with Sne Ia, the best fit yields $w_0$ very close to $0$ and Fig.~\ref{fig:SN.1} shows that a phantom transition is not excluded but unlikely. The fit with Standard Rulers leads to a similar result as well as when combining both datasets, as shown in Fig.~\ref{fig:C.1}, where specially the $w_0-\Omega_m^0$ contour plot favors $w_0\sim 0$. Nevertheless, the best fit for $w_1$ shows that the EoS parameter tends to an effective cosmological constant in the past ($z \gg 0$) whereas ends up slightly above the phantom barrier at small redshifts. Regarding the second parametrization (\ref{2.2}), the $w_0$ parameter is always negative at the best fit independently of the data source used, which gives rise to an EoS parameter $w_2$ that crosses the phantom barrier at large redshifts $z\gg z_0$ for the best fit, whereas remains above the phantom barrier at small redshifts. In addition, $w_0\sim 0$ is also favored, as shown in the right panels of Figs.~\ref{fig:SN.1}-\ref{fig:SR.1} and Fig.~\ref{fig:C.2}, but a phantom transition is within the confidence region of the contour plots, specifically the best fit leads to a phantom dark energy fluid in the past that tends to a non-phantom regime at small redshifts as pointed out above.\\ On the other hand, it is remarkable that neither models leads to future singularities at the best fit, as shown in Table \ref{table5}, although the possibility of the occurrence of a future singularity is not excluded within the confidence region of the contour plots, as depicted in Figs.~\ref{fig:SN.1}-\ref{fig:C.2}.\\ Hence, the analysis of the present manuscript shows that dealing with effective descriptions of the dark energy EoS can be well connected with the reconstruction of the underlying theory. In addition, both new parametrizations studied here show that a phantom epoch is compatible with the observational data in spite of that $\Lambda$CDM model is still very likely, although the best fit deviates a little bit from a cosmological constant. Moreover, the analysis shows that a singularity in the future is not excluded but also unlikely in both parametrizations. | 14 | 4 | 1404.3665 |
1404 | 1404.6795_arXiv.txt | We present the first J-band spectrum of Mrk 231, which reveals a large \ion{He}{1}*$\lambda 10830$ broad absorption line with a profile similar to that of the well-known \ion{Na}{1} broad absorption line. Combining this spectrum with optical and UV spectra from the literature, we show that the unusual reddening noted by \citet{veilleux13} is explained by a reddening curve like those previously used to explain low values of total-to-selective extinction in SNe Ia. The nuclear starburst may be the origin and location of the dust. { Spatially-resolved emission in the broad absorption line trough suggests nearly full coverage of the continuum emission region.} The broad absorption lines reveal higher velocities in the \ion{He}{1}* lines (produced in the quasar-photoionized \ion{H}{2} region) compared with the \ion{Na}{1} and \ion{Ca}{2} lines (produced in the corresponding partially-ionized zone). {\it Cloudy} simulations show that a density increase is required between the \ion{H}{2} and partially-ionized zones to produce ionic column densities consistent with the optical and IR absorption line measurements and limits, and that the absorber lies $\sim 100\rm \, pc$ from the central engine. These results suggest that the \ion{He}{1}* lines are produced in an ordinary quasar BAL wind that impacts upon, compresses, and accelerates the nuclear starburst's dusty effluent (feedback in action), and the \ion{Ca}{2} and \ion{Na}{1} lines are produced in this dusty accelerated gas. This unusual circumstance explains the rarity of \ion{Na}{1} absorption lines; without the compression along our line of sight, Mrk~231 would appear as an ordinary FeLoBAL. | } Mrk~231 is a nearby ($z=0.0421$) ultraluminous infrared galaxy that has a Seyfert 1 optical spectrum \citep{sanders88}. The infrared emission is thought to be a combination of AGN and starburst activity \citep[e.g.,][and references therein]{farrah03}. Recently, attention has been again drawn to this galaxy as a consequence of the discovery of a powerful, wide-angle, kiloparsec-scale molecular outflow \citep{rupke11}. Mrk 231's optical spectrum shows extreme \ion{Fe}{2} emission, undetected [\ion{O}{3}]$\lambda\lambda 4959,5007$, and somewhat broad and prominent Balmer lines. The most remarkable feature is the very strong broad \ion{Na}{1}D absorption line with $v \approx -4,500 \rm \, km\, s^{-1}$ \citep{aw72,boksenberg77, rudy85, sm85, hk87, boroson91, kdh92, forster95,rupke02, gallagher05, lipari05, rupke05, rodriguez09, veilleux13}, not to be confused with the few-hundred $\rm km\, s^{-1}$ \ion{Na}{1}D absorption consistent with the velocity of the molecular outflow \citep{rupke05}. While low-velocity \ion{Na}{1}D absorption from molecular outflows is relatively common \citep[e.g.,][]{rupke05}, high-velocity, broad absorption \ion{Na}{1}D lines are rare. The \ion{Na}{1}D absorption is accompanied by \ion{Ca}{2}$\lambda 3935,3970$ and \ion{He}{1}*$\lambda 3889$ in the optical bandpass \citep[e.g.,][]{aw72,boksenberg77,rudy85,sm85,hn87, boroson91,rupke02,lipari05,rodriguez09, veilleux13}. Near-UV and UV broad absorption lines were reported by \citet{smith95} and \citet{gallagher05}. { The broad absorption lines classify Mrk~231 as a broad absorption line Quasar (BALQSO). Broad absorption line quasars constitute between 10 and 40\% of quasars, depending on selection criteria \citep[e.g.,][]{hw03, trump06, dai08}.} Mrk~231 also exhibits \ion{Fe}{2} broad absorption lines, making Mrk~231 the nearest iron low-ionization broad absorption line quasar (FeLoBAL). { FeLoBALs are much rarer than BALQSOs, constituting only $\sim 2$\% of quasars, and again, the rate of incidence measured depends on the sample selection \citep{urrutia09, dai12}. BALQSOs have drawn intense interest in recent years, as their outflows may be key for understanding AGN feedback, i.e., the process by which an AGN influences its host galaxy, specifically, how does the outflow act to supress specifically, how it acts to suppress the rate of star formation in the host; see, for example, \citet{farrah12} who provides a discusison of the problem, and also has recently found evidence for an anticorrelation between the strength of the broad absorption-line outflow and the fractional contribution from star formation to the IR luminosity in a sample of FeLoBALs; they interpret this result as evidence for the AGN outflow curtailing star formation in the host galaxy.} As noted by \citet{veilleux13}, the presence of \ion{He}{1}*$\lambda 3889$ and \ion{Na}{1} in the same outflow appears to be problematic from a photoionization point of view. Metastable \ion{He}{1}* is formed by recombination onto He$^+$, and so it occurs in the \ion{H}{2} region of the outflow, in roughly the same gas that would generate \ion{C}{4} \citep[e.g.,][]{leighly11}. In contrast, neutral sodium has an ionization potential of only $5.14\rm \, eV$, i.e., less than that of hydrogen, so must occur in the partially-ionized zone beyond the hydrogen ionization front. It is distinguished from other low-ionization absorption lines such as \ion{Fe}{2} that are formed in the partially-ionized zone, however. As shown in \citet{lucy14}, copious \ion{Fe}{2} is produced just past the hydrogen ionization front; \ion{Na}{1}D is produced in basically neutral gas, much deeper in the slab, and farther from the illuminated face. Thus, \ion{He}{1}* and \ion{Na}{1} cannot be produced in gas with even close to the same ionization state, and so it is not necessarily easy to see how they can exhibit the same dynamics. Mrk~231 also has an unusual optical/UV continuum spectrum \citep{smith95, veilleux13}. Specifically, while the optical spectrum shows only modest reddening, the spectrum falls steeply in the near UV, and levels out shortward of $\sim 2400$\AA\/. Anomalously steep reddening has been seen in several BALQSOs \citep{hall02, leighly09, jiang13}, but the mechanism for this reddening remains unexplained. In the paper, we present new infrared spectra from Mrk~231. While Mrk~231 has been observed often in the infrared, a careful search of the literature revealed no previously published J-band spectra. Our spectrum reveals, for the first time, the broad \ion{He}{1}*$\lambda 10830$ absorption line (\S\ref{irtf}). We also find evidence for the appearance of a new \ion{He}{1}*$\lambda 10830$ velocity component at $11,520 \rm km \, s^{-1}$ (\S\ref{mdm}, \S\ref{apo}). We present new, high signal-to-noise-ratio blue optical spectra (\S\ref{kpno}). Combining our data with published spectra (\S\ref{digitize}), we present an analysis of the reddening in this object, finding that the unusual spectral shape is consistent with so-called ``circumstellar reddening'' (to be defined and discussed in \S\ref{goobar}). \S\ref{abs_lines} presents an analysis of the spectra, and extraction of the \ion{He}{1}* and \ion{Ca}{2} line profiles, which we compare with a \ion{Na}{1} profile obtained from the literature. After introducing a physical scenario for the absorption lines in \S\ref{picture}, we present in \S\ref{sims} a simple photoionization model, using {\it Cloudy}, that yields the ionic column densities required to produce the observed absorption lines. We systematically investigate the assumptions of our model, within the limitations of {\it Cloudy} in \S\ref{beyond} and provide an optimal model in \S\ref{optimize}. The results are summarized in \S\ref{summary}. \begin{deluxetable}{ccc} \scriptsize \tablewidth{0pt} \tablecaption{IRTF Spectrum} \tablehead{ \colhead{Observed Wavelength} & \colhead{Flux Density\tablenotemark{a}} & \colhead{Flux Density Error\tablenotemark{a}} \\ \colhead{($\mu$m)} & \colhead{($10^{-17}\rm \, erg\, s^{-1}\, cm^{-2}\, $\AA\/$^{-1}$)} & \colhead{($10^{-17}\rm \, erg\, s^{-1}\, cm^{-2}\, $\AA\/$^{-1}$)}} \startdata 0.805413 & 988.822 & 10.4281 \\ 0.805616 & 994.342 & 10.8305 \\ 0.805818 & 988.422 & 10.5220 \\ 0.806020 & 987.099 & 10.8525 \\ 0.806223 & 971.811 & 10.6738 \\ \enddata \tablecomments{Table \ref{irtf_spec_table} is published in its entirety in the electronic edition of the {\it Astrophysical Journal}. A portion is shown here for guidance regarding its form and content.} \tablenotetext{a}{Corrected for Galactic extinction.} \label{irtf_spec_table} \end{deluxetable} \normalsize | } We present the first J-band infrared, and new blue optical spectra of Mrk 231. Combining these with spectra taken from the literature, we discovered a physical solution for the unusual reddening in this object. {\it Cloudy} modeling revealed unusual physical conditions required to produce both \ion{He}{1}* and \ion{Na}{1} absorption lines in the same gas, and inspired a physical interpretation involving an interaction of a quasar BAL wind with gas ejected from a nuclear starburst. The specific results are as follows: \begin{itemize} \item We fit the broad band spectrum from $\sim 2300$ \AA\/ to $2.3\,\rm \mu m$ with a continuum model, a low-temperature black body, and a circumstellar reddening and extinction curve originally proposed to explain the low values of observed $R_V$ in SNe Ia \citep{goobar08}. Circumstellar reddening is distinguished by large optical depths, approaching 1, producing increased extinction in the blue \& UV (due to longer scattering path lengths), along with light scattered back into the line of sight as a secondary effect. We obtained an excellent fit using extinction-shape parameters similar to those for Milky Way dust (\S\ref{goobar}). We suggested that the dust is produced in the nuclear starburst lying $\sim 100\rm \, pc$ from the nucleus \citep[e.g.,][]{davies04}. \item The infrared spectrum revealed a deep \ion{He}{1}* absorption line that is very similar in profile to the well-known \ion{Na}{1} absorption line. Two newer infrared spectra revealed evidence for the appearance of a new absorption line component near $11,000 \rm \, km\, s^{-1}$. We modeled the optical and infrared spectra to extract the line profiles (\S\ref{nuclear}, \S\ref{hei10830}). After accounting for the $\sim 100\rm \, Myr$ nuclear starburst contribution, we inferred that the absorber essentially completely covers the quasar continuum emission region. The absorption line profiles indicated that the lines produced in the \ion{H}{2} region, including \ion{He}{1}*$\lambda 10830$ and \ion{He}{1}$\lambda 3889$, have a higher velocity than the low-ionization lines produced in the partially-ionized and neutral gas, including \ion{Ca}{2} and \ion{Na}{1} (\S\ref{abs_lines_2}). \item {\it Cloudy} modeling showed that in order to produce both the \ion{He}{1}* and the \ion{Na}{1} absorption lines, a density increase is required between the \ion{H}{2} region, which produces the \ion{He}{1}* lines, and the partially-ionized/neutral region, the origin of the \ion{Ca}{2} and \ion{Na}{1} lines. We first modeled this effect as a constant pressure gas (\S\ref{cloudy}). The models are able to produce the measured column densities and limits if the gas lies $\sim 100\rm \, pc$ from the central engine, i.e., in the vicinity of the nuclear starburst \citep[e.\ g.,][]{davies04}. These facts, along with the velocity differences of the lines, and the inferred full covering, led us to a physical scenario in which the \ion{He}{1}* absorption arises in a quasar BAL outflow that impacts and compresses dusty gas originating in the starburst, and this swept up gas is the origin of the \ion{Na}{1} and \ion{Ca}{2} lines (\S\ref{picture}). In addition, we noted that just such an interaction between a quasar outflow and surrounding star-forming gas may be an example of quasar feedback, thought to be necessary to shut down star formation during the co-evolution of black holes and quasars. \item In \S\ref{beyond} we examined the effects of modifying the assumptions made in our initial {\it Cloudy} model. In addition, the outflow masses and kinetic luminosities inferred were very large if the global covering fraction is $\Omega=0.2$. Constant pressure gas may not be a reasonable assumption for a $\sim -4,500\rm \, km\, s^{-1}$ outflow, and therefore we instead experimented with a density increase between the \ion{H}{2} region and the partially-ionized zone that might be physically realized in a shock, and found that a density increase works as well as constant pressure. We discovered that the area of parameter space producing models consistent with the measurements is broadened when we consider abundances enhanced by stellar processing in the starburst, and when the spectral energy distribution is soft. We also suggested that the \ion{H}{2} BAL outflow is free of dust, while the partially-ionized zone may have circumstellar dust. This situation, which could further reduce the inferred column density and therefore kinetic luminosity, cannot presently be modeled using {\it Cloudy} (which cannot model variable dust properties as a function of depth). \item Finally, in \S\ref{optimize}, we presented full grids using the optimal parameters obtained in \S\ref{beyond}. For a soft spectral energy distribution and enhanced abundances, a density enhancement by a factor of only four can still produce the lines we observe, although over a limited region of parameter space. A density enhancement by a factor of 25 opened up parameter space considerably. A shock would be expected to produce a density contrast of 4 for gas with $\gamma=5/3$, and much larger density contrasts are inferred in supernovae, depending on the density profiles of the outflow and circumnuclear gas. \item Most of the simulations producing the observed ionic column densities favor low densities, i.e., $\log n \approx 4$. This density is characteristic of the narrow-line region in AGN. We find that, assuming a global covering fraction of $\Omega = 0.2$, the inferred equivalent width of the [\ion{O}{3}] emission line should be 100--200\AA\/. Such a huge line is not seen in Mrk~231. We note that the line may not necessarily be predicted to be sharp, but rather could be smeared by a range of line-of-sight velocities, or obscured by dust. \end{itemize} While we consider the question of the cause of the anomalous reddening essentially solved in this paper, a number of questions regarding the absorption lines remain. Some may be addressed by the upcoming {\it HST} observations of Mrk~231. For example, additional UV and near-UV absorption line measurements may be able to further refine the {\it Cloudy} modeling; this may be complicated by differential continuum and starburst covering fractions and extinctions. Dynamical modeling may be able to determine whether shocks could produce density jumps as large as the simulations require. Further development of {\it Cloudy} may ultimately allow specification of dust properties as a function of depth into the gas slab. In addition, our models are 1-D, and so cannot explicitly take into account the effect of the circumstellar reddening on the photoionization results. At any rate, the unusual set of circumstances required to produce the observed optical and infrared absorption lines may explain why \ion{Na}{1} lines are so rare, or at least why the line in Mrk~231 is so exceptional; without the interaction of the BAL wind and starburst gas along our line of sight, Mrk~231 might instead look like an ordinary FeLoBAL quasar. | 14 | 4 | 1404.6795 |
1404 | 1404.1954.txt | {Debris disks have been found primarily around intermediate and solar mass stars (spectral types A-K), but rarely around low-mass M-type stars. This scarcity of detections in M~star surveys can be confronted with the predictions of the steady state collisional evolution model. First, we determine the parameters of the disk population evolved with this model and fit to the distribution of the fractional dust luminosities measured in the surveys of A- and FGK-type stars observed by the infrared satellite {\it Spitzer}. Thus, in our approach, we stipulate that the initial disk mass distribution is bimodal and that only high-mass collisionally-dominated disks are detected. The best determined parameter is the diameter $D_c$ of the largest planetesimals in the collisional cascade of the model, which ranges between 2 and 60~km, consistently for disks around A- and FGK-type stars. Second, we assume that the same disk population surrounds the M~dwarfs that have been the subjects of debris disk searches in the far-infrared with {\it Spitzer} and at submillimeter wavelengths with radiotelescopes. We find, in the framework of our study, that this disk population, which has been fit to the AFGK data, is still consistent with the observed lack of disks around M~dwarfs with {\it Spitzer}. } | In our study, we have fit the steady state collisional evolution model of \citet{Wyat07a} to the fractional dust luminosity distributions measured with {\it Spitzer} for the A and FGK stars. Hence, we have postulated that these distrubutions correspond to the population of high-mass collisionally-dominated disks that are detectable in the current surveys and that can be analyzed {\it per se}. There is a distinct population of low-mass disks that are not considered in the study because they are undetectable. The results of our study must be understood in this framework. The model parametrized with $M_{mid}, D_c, R1$, and $R2$ satisfactorily fits the observations and the two resulting best-fit regions for the samples of A and FGK stars are overlapping, indicating that there are solutions common to both samples. We have applied this steady state collisional evolution model to potential high-mass collisionally-dominated disks around M~dwarfs surveyed prior to the {\it Herschel} satellite in the far-infrared by \citet{Gaut07} and in the submillimeter domain by \citet{Lest06, Lest09}. Within the framework of this study, we have shown that the disk population fit to the AFGK stars is still consistent with the lack of disks detected around M-dwarfs in the far-infrared and submillimeter surveys. In a future study, we shall apply our novel approach to study evolution of debris disks in the larger sample of M~dwarfs in the unbiased Herschel DEBRIS survey. | 14 | 4 | 1404.1954 |
|
1404 | 1404.3515_arXiv.txt | Our analysis in Papers I and II (Grechnev \textit{et al.}, 2014, Solar Phys. 289, 289 and 1279) of the 18 November 2003 solar event responsible for the 20 November geomagnetic superstorm has revealed a complex chain of eruptions. In particular, the eruptive filament encountered a topological discontinuity located near the solar disk center at a height of about 100 Mm, bifurcated, and transformed into a large cloud, which did not leave the Sun. Concurrently, an additional CME presumably erupted close to the bifurcation region. The conjectures about the responsibility of this compact CME for the superstorm and its disconnection from the Sun are confirmed in Paper~IV (Grechnev \textit{et al.}, Solar Phys., submitted), which concludes about its probable spheromak-like structure. The present paper confirms the presence of a magnetic null point near the bifurcation region and addresses the origin of the magnetic helicity of the interplanetary magnetic clouds and their connection to the Sun. We find that the orientation of a magnetic dipole constituted by dimmed regions with the opposite magnetic polarities away from the parent active region corresponded to the direction of the axial field in the magnetic cloud, while the pre-eruptive filament mismatched it. To combine all of the listed findings, we come to an intrinsically three-dimensional scheme, in which a spheromak-like eruption originates \textit{via} the interaction of the initially unconnected magnetic fluxes of the eruptive filament and pre-existing ones in the corona. Through a chain of magnetic reconnections their positive mutual helicity was transformed into the self-helicity of the spheromak-like magnetic cloud. | \label{S-introduction} An extreme geomagnetic storm on 20 November 2003 was caused by the interaction of the Earth's magnetosphere with an interplanetary magnetic cloud (MC), whose magnetic helicity, $H_\mathrm{m}$, was positive. The very strong magnetic field in the MC of up to $|B|_{\max} \approx 56$~nT had a long-lasting negative $B_z$ component (up to $B_{z\, \max} \approx -46$~nT). These circumstances were crucial in identifying the solar source for the MC. \inlinecite{Gopal2005} and \inlinecite{Yurchyshyn2005} definitely associated the MC with the filament eruption in the active region (AR) 10501 (we will use henceforth the last three digits for brevity) on 18 November. These authors considered the direction of the axial magnetic field in the pre-eruptive filament to correspond to the expected projection $B_z < 0$. With such a direction of the axial field, the current helicity of the filament, $H_\mathrm{c}$, is positive. From the condition $\mathrm{sign} (H_\mathrm{m}) = \mathrm{sign} (H_\mathrm{c})$, which is valid for a linear force-free magnetic field, it follows that the magnetic helicity is also positive. Later \inlinecite{Moestl2008} found a correspondence between the flare reconnected magnetic flux, measured as the flare ribbon flux, and the poloidal magnetic flux of the MC under the assumption that the MC was a part of a magnetic flux rope with a length of 0.5--2~AU. The studies by \inlinecite{Gopal2005}, \inlinecite{Yurchyshyn2005}, and \inlinecite{Srivastava2009} gave the impression of an acceptable correspondence between the conditions of the eruption in AR~501 and the parameters of the MC observed in the Earth orbit: $H_\mathrm{m} > 0, B_z < 0$. The study of \inlinecite{Chandra2010} changed this situation. From the observed morphological features they found that the large-scale magnetic field in AR~501 had a negative helicity sign. This finding seemingly contradicted what was expected from the magnetic helicity conservation requiring the same sign of the magnetic helicity in the AR and MC. This circumstance has raised a question, why the AR, which had a global negative magnetic helicity, could expel a positive-helicity MC to the interplanetary medium. One possible answer was proposed by \inlinecite{Chandra2010}, who found a localized positive helicity injection in the southern part of AR~501 and concluded that the right handedness of the observed MC was due to the ejection from this portion of the AR. On the other hand, \inlinecite{Leamon2004} in their study of 12 interplanetary MCs and related solar active regions have found: \textit{i})~a significant difference between the total twist of the magnetic field inside active regions, $(\alpha L)_\mathrm{AR}$, and that in the MC, $(\alpha L)_\mathrm{MC}$; \textit{ii})~the absence of any significant sign relationship between them. The authors used the linear force-free approximation, $\alpha$ is a constant. The dipole scale, $L_\mathrm{AR}$, was measured as the distance between the centroids of the positive and negative fluxes in the magnetogram of an AR. The magnetic field in an MC was fit with the Lundquist magnetic model with $L_\mathrm{MC} = 2.5$~AU length. Findings (\textit{i}) and (\textit{ii}) have compelled \inlinecite{Leamon2004} to conclude that ``\textit{magnetic clouds associated with active region eruptions are formed by magnetic reconnection between these regions and their larger-scale surroundings, rather than simple eruption of preexisting structures in the corona or chromosphere}''. \inlinecite{ZhangLow2003} described a similar phenomenon analytically by using the idealized example of the axially-symmetric reconfiguration of two twisted magnetic fluxes from their unconnected initial state to the connected relaxed state. They have shown that magnetic reconnection can reverse the twist direction of a flux rope emerging into preexisting fields under the conservation of the total relative magnetic helicity. The complex reconnection of a flux rope with the adjacent field in complex magnetic topology has been also described by, \textit{e.g.}, \inlinecite{Lugaz2011}, \inlinecite{Zuccarello2012}, and \inlinecite{Masson2013}. \citeauthor{Grechnev2008} (\citeyear{Grechnev2008,Grechnev2011_AE,Grechnev2013_anomal}) have found observational evidence of magnetic reconnection between the internal field belonging to the eruptive filament and the preexisting coronal field. This is a phenomenon of plasma dispersal from an eruptive filament over the solar surface that is visible as the disintegration of the filament. The whole mass of an eruptive filament or a considerable fraction of its mass does not leave the Sun as a part of a CME. The motion of the cool plasma of the eruptive filament continues along new magnetic field lines passing inside the eruptive filament and ending far on the solar surface. Clouds of such plasma can screen the emission of compact sources in active regions as well as the emission from quiet solar areas. Absorption phenomena can be observed in microwaves and also in the He~\textsc{ii} 304~\AA\ line. Events of such a kind are associated with active region eruptions. They have been rarely detected for observational reasons. The analysis of the solar geoeffective event of 18 November 2003 by \inlinecite{Grechnev2014_I} (hereafter Paper~I) and \inlinecite{Grechnev2014_II} (hereafter Paper~II) has revealed that the major eruption in this event, \textit{i.e.}, the eruption of the U-shaped filament, which we call F1, from AR~501, was also not a simple one. The eruptive filament bifurcated and transformed into a large Y-shaped cloud, which moved from the region of bifurcation (Rb) to the South--West across the solar disk toward the limb. Figure~\ref{F-shape_transform} illustrates what has happened to the eruptive filament, as shown by the H$\alpha$ images produced by the Kanzelh{\"o}he Solar Observatory (KSO), the \textit{Extreme-ultraviolet Imaging Telescope} (EIT; \opencite{Delab1995}), on board the \textit {Solar and Heliospheric Observatory} (SOHO), and the \textit{Spectroheliographic X-ray Imaging Telescope} (SPIRIT; \inlinecite{Zhitnik2002} and \inlinecite{Slemzin2005}), on board the \textit{Complex Orbital near-Earth Observations of Activity of the Sun} (CORONAS-F) satellite \cite{{OraevskySobelman2002},{Oraevsky2003}}. \begin{figure} % \centerline{\includegraphics[width=\textwidth] {shape_transform.eps} } \caption{Bifurcation of the main eruptive filament F1. (a,b)~Pre-eruption H$\alpha$ line-center image (KSO). The frames denote the fields of view of the four images shown in the lower row (b--e). The axes are in arc seconds from the solar disk center. The turquoise oval in all the images denotes the region Rb where the filament bifurcated. The yellow curves roughly outline the frontal edge of the filament before and during the eruption. (c)~SOHO/EIT image in the 195~\AA\ channel. (d)~KSO H$\alpha$ image observed in the far blue wing. (e)~Y-like cloud in the CORONAS-F/SPIRIT 304~\AA\ image.} \label{F-shape_transform} \end{figure} The masses of the Y-like cloud and the pre-eruption filament were similar (Paper~I); on the other hand, remnants of the filament were not evident in the southwestern CME observed by the \textit{Large Angle and Spectrometric Coronagraph} (LASCO) that was previously regarded as the source of the 20 November geomagnetic storm. The observations analyzed in Paper~I and Paper~II suggest a possible additional eruption in the interval from 08:07 to 08:14~UT above the bifurcation region close to the solar disk center that could be the source of the interplanetary MC on 20 November. These facts disfavor the simple scenario, in which the 20 November MC is considered as a flux rope formed directly from a structure initially associated with the pre-eruptive filament F1 in region 501. As mentioned, the right-handed MC produced in this event and responsible for the superstorm had a very strong magnetic field near Earth of up to $|B|_{\max} \approx 56$~nT and $B_{z\, \max} \approx -46$~nT. According to \inlinecite{Moestl2008}, its inclination to the ecliptic plane was $\theta = -(49-87)^{\circ}$, and the magnetic flux in this plane was $0.55 \times 10^{21}$~Mx; however, its significant part could be lost by reconnection in the interplanetary space. In \inlinecite{Grechnev2014_IV} (hereafter Paper~IV), we additionally find that the MC was compact, with a size of about 0.2~AU, and had some atypical properties, such as a wide range of proton temperatures, from $2 \times 10^4$~K to $3\times 10^5$~K; its magnetic structure was closed, disconnected from the Sun, and probably had a spheromak configuration. The present paper (hereafter Paper~III) endeavors to understand how the catastrophe of the eruptive filament could occur and create the right-handed spheromak-like MC. Section~\ref{S-outline} outlines the eruptive event and results of its analysis. Section~\ref{S-helicity} analyzes the helicity in AR~501. In Section~\ref{S-large_scale_config} we address the causes of the bifurcation of the eruptive filament. In Section~\ref{S-mc_formation} we try to understand how the MC could be formed. Section~\ref{S-summary} briefly summarizes the outcome of this study. | \label{S-summary} The scenario of the 18 November 2003 event does not correspond to the concept of a simple eruption directly from AR~501, in which the twist helicity of an eruptive structure or active region determines the handedness of the interplanetary magnetic cloud. The NLFF extrapolation of AR~501 shows an excess of negative twist, which is opposite to the positive sign of twist in the MC. To solve this contradiction, we have used the positive mutual helicity between the pre-eruptive filament and the flux tubes of a magnetic domain of the large-scale quadrupole configuration. The interaction of these magnetic fluxes presumably occurred as the eruptive filament passed in the neighborhood of the coronal magnetic null point. The positive mutual helicity of these two fluxes changed through magnetic reconnections into the positive self-helicity of a spheromak-like structure, whose geometry and parameters correspond to the magnetic cloud, which reached Earth. In Paper~IV, we analyze the interplanetary disturbance responsible for the 20 November superstorm and outline the overall scenario of the whole event. \begin{acks} We thank V.~Yurchyshyn, who kindly supplied the vector magnetogram of AR~501 observed at BBSO, and M.~Temmer for the H$\alpha$ data. We thank the co-authors of our Papers I, II, and IV, who are not involved in this study. We appreciate an anonymous reviewer for valuable remarks and comments. We are grateful to the instrumental teams of the Kanzelh{\"o}he Solar Observatory; TRACE and CORONAS-F missions; MDI and EIT on SOHO (ESA and NASA) for the data used here. This study was supported by the Russian Foundation of Basic Research under grants 11-02-00757, 11-02-01079, 12-02-00008, 12-02-92692, and 12-02-00037, and the Ministry of Education and Science of Russian Federation, projects 8407 and 14.518.11.7047. \end{acks} | 14 | 4 | 1404.3515 |
1404 | 1404.3209_arXiv.txt | The current velocity of the Smith Cloud indicates that it has undergone at least one passage of the Galactic disc. Using hydrodynamic simulations we examine the present day structure of the Smith Cloud. We find that a dark matter supported cloud is able to reproduce the observed present day neutral hydrogen mass, column density distribution and morphology. In this case the dark matter halo becomes elongated, owing to the tidal interaction with the Galactic disc. Clouds in models neglecting dark matter confinement are destroyed upon disc passage, unless the initial cloud mass is well in excess of what is observed today. We then determine integrated flux upper limits to the gamma-ray emission around such a hypothesised dark matter core in the Smith Cloud. No statistically significant core or extended gamma-ray emission are detected down to a 95\% confidence level upper limit of 1.4 $\times 10^{-10}$ ph cm$^{-2}$ s$^{-1}$ in the 1--300 GeV energy range. For the derived distance of 12.4 kpc, the {\it Fermi} upper limits set the first tentative constraints on the dark matter cross sections annihilating into $\tau^{+}{\tau}^{-}$ and $b\bar{b}$ for a high-velocity cloud. | The mapping of dark matter substructure around the Galaxy is crucial to understanding how our own Milky Way was assembled over cosmic time. Numerical simulations of the concordant cosmology, dark-energy and cold dark-matter ($\Lambda$CDM), predict a multitude of dark-matter subhaloes going down in mass to approximately $10^{-4}~M_\odot$ \citep[\eg][]{Klypin1999,Moore1999,Diemand2007,Springel2008}. On the detectable scale of dwarf galaxies, this discrepancy approaches an order of magnitude between the observed dwarf galaxies and the number of predicted subhaloes \citep{Mateo1998,Weinberg2013}. Recent discoveries of ultra-faint dwarf galaxies go someway towards filling this gap \citep[\eg][]{Willman2005,Belokurov2007}. Unfortunately, due to the minuscule stellar populations, finding ultra faint dwarfs continues to be challenging and the exact number of them are unknown. The low stellar content of them however, makes them potentially excellent objects to search for dark matter annihilation signals \citep{Charbonnier2011}. So far, no significant detections have been made, although upper limits have been placed on the properties of dark matter \citep[\eg][]{dwarfs,Natarajan2013,Fermi2013}. In addition to dwarf galaxies, the Galactic halo contains a large number of intriguing \HI{} substructures in the form of high velocity clouds (HVCs) \citep{Wakker1997}. The suggestion of dark matter surrounding a population of these HVCs has been around for over a decade \citep[\eg][]{Blitz1999,Quilis2001}. Subsequent investigation has determined that any such population is likely to be small in comparison to dark matter free HVCs, both around our own Galaxy \citep{Saul2012} and also around other galaxies \citep{Chynoweth2011}. Near the disc some of the HVCs are likely to arise through a galactic fountain, where numerous supernova launch gas from the disc of the Galaxy \citep{Bregman1980}, however, this process is likely to produce small, intermediate velocity clouds over HVCs \citep{Ford2010}. Many HVCs that lie within 50 kpc of the Milky Way likely had their origin in the Magellanic Stream and thus arose from the LMC and SMC \citep{Putman2004}. Extragalactic HVCs are seen at a wide range of projected distances, often exceeding $150$~kpc from the nearest galaxy, are likely to be clumps of pristine gas infalling for the first time and possess a phase-space distribution that is incompatible with the expected dark matter substructure \citep{Chynoweth2011}. Even at low projected distances ($<$$50$~kpc) many of the extragalactic HVCs are not associated with regions of star formation suggesting that they too are infalling for the first time \citep{Thilker2004,Westmeier2008}. Despite the abundance of other sources for HVCs, explaining why all dark matter subhaloes that failed to form stars lack any gas is difficult and a small fraction of HVCs may be such objects. In order to investigate the dark matter content in \HI{} clouds, we recently searched for gamma-ray emission from dark matter annihilation at the location of several compact \HI{} clouds in the GALFA-HI Compact Cloud Catalogue \citep{Mirabal2013}. However, due to poorly constrained dark matter profiles and unknown distances to these objects severely limited our analysis. To advance more thoroughly, we need to examine systems with better distance determinations and are likely to be candidates for dark matter embedded HVCs. The best candidate for such a search is the Smith Cloud, a massive \HI{} system near the Galactic disc \citep{Smith1963}. The Smith Cloud is located close by at $12.4\pm1.3$~kpc \citep{Lockman2008}, and of particular appeal is that a dark-matter subhalo seems to be required for the survival of the gas cloud after a passage through the Galactic disc \citep{Nichols2009}. Given its relative proximity, apparent orbit and large mass it appears to qualify as the ideal astrophysical site to test the dark matter confinement of \HI{} clouds. Here we search for potential gamma-ray emission from the dark matter annihilation of weakly interacting particles (WIMPs) around the Smith cloud. In Section 2 we present arguments in favour of a dark-matter subhalo surrounding the Smith cloud. Next, we present the {\it Fermi}-LAT analysis and derive {\it Fermi} upper limits to the gamma-ray flux. In Section 5, we turn the results into limits on annihilation cross sections. Finally, we summarise our results and present our conclusions. | \label{interp} We have undertaken analytic calculations and numeric simulations to investigate whether the Smith Cloud is encapsulated within a dark matter halo. In the absence of dark matter, only the most dense and correspondingly massive clouds survive the passage through the disc of the Galaxy. Such massive objects show column densities much higher than the Smith Cloud is observed to have, and morphologically consist of a dense core with connected tendrils of gas. With a dark matter halo, a HVC is able to survive the passage through substantially intact. At low densities, $n\sim0.2$~cm$^{-3}$, the resulting cloud contains a column density similar to that of the Smith Cloud and a similar morphological structure to the Smith Cloud of fragmented clumps unattached to the main structure but following the orbit. Such simulations suggest that the idea of the Smith Cloud being encapsulated by a dark matter halo remains plausible, in particular for this orbit, morphology and $N_{\rm HI}$ distribution. For more massive clouds, the inner densities are approaching that of star forming regions in metal poor dwarf galaxies \citep{Ekta2008}. The spatial segregation of gas free dSphs and dIrrs, where the former is mostly found close to the main galaxy, in the Milky Way and Andromeda is evidence for efficient gas stripping closer to the halo centres \citep{Grcevich2009}. Any star formation would lead to supernova feedback greatly assisting the stripping of gas from the HVC \citep{Gatto2013}, regardless of the presence of dark matter. In light of \cite{Gatto2013}, the absence of stellar feedback in the Smith Cloud makes the fact that we see such a gas rich structure, which possibly is an "unformed" dwarf galaxy, less of a timing problem, as the compact gas cloud may survive many disc crossings without being completely destroyed via ram pressure/tidal stripping. To fully understand the plausibility of relating an object like the Smith Cloud to non-star forming substructure, a fully cosmological context is necessary. If encapsulated by dark matter, the Smith Cloud therefore inhabits a narrow region between being too light to survive ram pressure removing it from the host dark matter subhalo and being too massive with resulting star formation greatly assisting this stripping. We report photon flux upper limits for the gamma-ray emission at the current position of the Smith Cloud using 4.85 years of accumulated {\it Fermi}-LAT data to investigate the properties of such a dark matter halo. We exclude WIMPs annihilating into $\tau^{+}{\tau}^{-}$ and $b\bar{b}$ final states down to $\langle\sigma v\rangle_{\chi} \sim 3\times 10^{-25}$ cm$^{3}$ s$^{-1}$ for masses around 10 GeV. The obvious caveat is that dark matter might have been shed during the history of the cloud, however, the simulations have the dark matter retain the cloud core up to its present day location with only minor elongation. To investigate this case we compare the counts from the projected orbit and anti-orbit. We find no evidence for excess gamma-ray emission along the predicted trajectory of the cloud system. In addition, there is no morphological gamma-ray structures overlapping the cometary structure reported in \HI{}. Despite this failed effort, gamma-ray observations still offers one of the few available opportunities to diagnose the dark matter content of gaseous clouds. Upcoming experiments such as the Cherenkov Telescope Array (CTA) will be able to achieve improved angular resolution and sensitivity around the Smith Cloud for energies above 100~GeV \citep{cta,doro}. The velocity averaged annihilation cross section upper bounds obtained around the Smith cloud are compatible with limits from other searches reported in the dwarf galaxies \citep{dwarfs} and galaxy clusters \citep{clusters}. Our results rest upon the assumption that there is indeed a dark-matter subhalo seeding the Smith Cloud. To this extent we undertook simulations demonstrating that such an assumption is plausible given the disc passage the Smith Cloud is likely to have undertaken. As well as demonstrating that the Smith Cloud's structure can be reproduced through a dark matter embedded HVC, such a result suggests that clouds with dark matter that have passed a disc may take on the comet like morphology observed in expected in DM free HVCs which have not yet crossed the disc \citep[\eg][]{Putman2011,Plockinger2012}. The latter still needs to be verified with future observations, however, it provides potentially the best candidate for a dark matter confined HVC. | 14 | 4 | 1404.3209 |
1404 | 1404.2808_arXiv.txt | CCD observations of 68 eclipsing binary systems, candidates for containing $\delta$~Scuti components, were obtained. Their light curves are analyzed using the Period04 software for possible pulsational behaviour. For the systems QY~Aql, CZ~Aqr, TY~Cap, WY~Cet, UW~Cyg, HL~Dra, HZ~Dra, AU~Lac, CL~Lyn and IO~UMa complete light curves were observed due to the detection of a pulsating component. All of them, except QY~Aql and IO~UMa, are analysed with modern astronomical softwares in order to determine their geometrical and pulsational characteristics. Spectroscopic observations of WY~Cet and UW~Cyg were used to estimate the spectral class of their primary components, while for HZ~Dra radial velocities of its primary were measured. O$-$C diagram analysis was performed for the cases showing peculiar orbital period variations, namely CZ~Aqr, TY~Cap, WY~Cet and UW~Cyg, with the aim of obtaining a comprehensive picture of these systems. An updated catalogue of 74 close binaries including a $\delta$~Scuti companion is presented. Moreover, a connection between orbital and pulsation periods, as well as a correlation between evolutionary status and dominant pulsation frequency for these systems is discussed. | In general, close binaries, and especially the eclipsing ones, are stellar objects whose absolute parameters and evolutionary status can be easily derived from observations. Single $\delta$~Scuti stars show discrepancy from binary-members regarding their evolutionary status. The single ones are situated on the Main Sequence (hereafter MS) or moving from it to the giant branch. On the other hand, the $\delta$~Scuti components in binaries show slow evolution through the MS, thus they are very useful tools to diagnose this extraordinary part of a stellar lifetime. This difference in evolution is connected with the mass transfer process and tidal distortions occurring in binary systems during their MS life \citep{MK03,SO06a}. Therefore, single and binary-contained $\delta$~Scuti stars, although they present similar pulsational properties, should not be considered of the same type due to a likely different evolutionary past. Especially the ones in classical Algols show variable pulsational characteristics due to mass gain \citep{MK04,MK07}, a process that is responsible for differences in the excitation mechanism, i.e. $\kappa$-driven oscillations compared to tidally induced or mass-accretion induced oscillations. Additionally, calculation of the absolute parameters and the identification of the oscillating characteristics of a binary's pulsational component provide the means to obtain a detailed picture of the star. Obviously, the larger the sample of such stars the more information and conclusions can be derived. During the last decade interest for pulsating stars in close binaries has increased significantly and a lot of discoveries announced. \citet{MK04} introduced the oEA (oscillating EA) stars as the (B)A-F spectral type mass-accreting MS pulsators in semi-detached Algol-type eclipsing binary (hereafter EBs) systems. \citet{SO06a} made a first attempt to find a connection between pulsation and orbital periods of systems with $\delta$~Scuti component and resulted into a linear relation. \citet{ZH10} published a catalogue containing 89 systems and distinguished them according to their pulsational properties. \citet{SO11} also published a similar list including 43 cases of systems including a $\delta$~Sct component. Space missions such as \textit{CoRot} and \textit{Kepler} have been discovering many pulsating stars in binaries (cf. \citealt{SOU11}; \citealt{DA10}). Their measurements provide the means to derive many pulsation frequencies and identify the oscillating modes with unprecedented accuracy. The present work is a continuation of the survey for $\delta$~Scuti components in EBs of \citet{LN09}. Eight new such systems are presented herein for the first time, while the results of some cases that were recently discovered by other investigators are confirmed. The majority of the observed systems were selected from the lists of \citet{SO06b}. The results of both surveys, which enriched significantly ($\sim14\%$) the sample of such binaries, can therefore be used for future space and/or ground based observing campaigns. The study of the `\textsl{O}bserved$-$\textsl{C}alculated' times of minima variation(s) (hereafter O$-$C analysis) of an EB provides valuable information for the mechanisms which form its orbital period (e.g. mass transfer, third body, magnetic influences etc). On the other hand, the `snapshot' of the binary, i.e. its light curve (hereafter LC), leads us to understand directly which physical processes are occurring (e.g. third light existence, Roche Lobe filling). The combined information from these two independent methods of analysis provides a more comprehensive view of a binary system. Eleven systems exhibiting oscillating behaviour were selected for further observations and eight of them are studied using efficient modern techniques for LC and photometric frequency analyses. The remaining three cases QY~Aql, BO~Her, and IO~UMa, will be presented in a future work. Four of these eleven were found to have orbital period changes from O$-$C analysis. \section[]{Observations and data reduction} The photometric observations were carried out during September 2008 - September 2011 at the Gerostathopoulion Observatory of the University of Athens (UoA) located in Athens, Hellas, and at the Kryonerion (Kry) Astronomical Station of the Astronomical Institute of the National Observatory of Athens located at Mt.~Kyllini, Corinthia, Hellas. The instrumentation used for the observations is described in detail in Table~1. Aperture photometry was applied to the data and differential magnitudes for all systems were obtained using the software \emph{MuniWin} v.1.1.26 \citep{HR98}. The adopted observational strategy in this survey was the same as that described in detail in the previous paper \citep{LN09}. Briefly, the observational guidelines were: i) the time span should be greater than 3~hr, ii) the filter $B$ or $V$ should be used, iii) the comparison star should be of similar magnitude and spectral type to the variable, iv) appropriate exposure times and binning modes must be used for the best possible photometric $S/N$ (signal-to-noise ratio). In Table~2 we list: the name of the system as given in the GCVS catalogue \citep{SA12}, its apparent magnitude ($m$) and spectral type (\emph{S.T.}) as given in SIMBAD, the filters (\emph{F}) used, the number of nights (\emph{N}) of observations, the total time span (\emph{T.S.}), the standard deviation (\emph{S.D.}) of the observed points (mean value), the phase intervals (\emph{P.I.}), the dominant pulsation frequency ($f_{\rm dom}$) found and its semi-amplitude ($A_{\rm B}$) in $B$-filter (see section~6 for details), and the abbreviation for the instrumentation (\emph{In}) used for each case according to Table~1. For the majority of the systems, the observations were obtained outside the primary eclipse, since the primary component (hotter one) was the candidate $\delta$~Sct star. The eight systems found to exhibit pulsational behaviour and whose complete LCs were obtained are listed in Table~3 along with the corresponding Comparison and Check stars. The spectroscopic observations were obtained with the 1.3~m Ritchey-Chr\'{e}tien telescope at Skinakas Observatory, Mt.~Ida, Crete, Hellas, on 6 October 2010 for WY~Cet, UW~Cyg and on 20 and 24 September 2011 for HZ~Dra. We used a 2000$\times$800 ISA SITe CCD camera attached to a focal reducer, with a 2400~lines/mm grating and slit of 80~$\mu$m. This arrangement yielded a nominal dispersion of $\sim$0.55~\AA/pixel and wavelength coverage between 4782-5864~\AA. Data reduction was performed using the \textsl{Radial Velocity reductions} v.2.1d software \citep{NE09}. The frames were bias subtracted, a flat field correction was applied, and the sky background was removed. The spectral region was selected so as to include H$_{\beta}$ and sufficient metallic lines. Before and after each on-target observation, an arc calibration exposure (NeHeAr) was recorded. \begin{table} \begin{minipage}{\columnwidth} \centering \caption{The instrumentation setups (Telescope \& CCD) used during the photometric observations, their location ($Loc.$) and abbreviation ($Ab.$).} \begin{tabular}{lcccc} \hline Telescope & CCD* & $Loc.$ & $Ab.$ \\ \hline Cassegrain - 0.4~m, $f$/8 & ST--8XMEI & UoA & $A1$ \\ Cassegrain - 0.4~m, $f$/8 & ST--10XME & UoA & $A2$ \\ Newt. Reflector - 0.2~m, $f$/5 & ST--8XMEI & UoA & $A3$ \\ Newt. Reflector - 0.25~m, $f$/4.7 & ST--8XMEI & UoA & $A4$ \\ Cassegrain - 1.2~m, $f$/13 & AP47p & Kry & $K$ \\ \hline \multicolumn{5}{l}{*The CCDs are equipped with the Bessell $U,~B,~V,~R,~I$}\\ \multicolumn{5}{l}{photometric filters} \end{tabular} \end{minipage} \end{table} \begin{table*} \centering \caption{The total log of observations.} \scalebox{0.94}{ \begin{tabular}{lcccccccccc} \hline System & $m$ & $S.T.$ & $F$ & $N$ & $T.S.$ & $S.D.$ & $P.I.$ &$f_{\rm dom}$&$A_{\rm B}$& $In$ \\ & (mag) & & & & (hrs) & (mmag) & & (c/d) & (mmag) & \\ \hline And CP & 11.80 ($B$) & A5 & $B$ & 1 & 6 & 5.8 & 0.20-0.26 & -- & -- & A2 \\ And TW & 9.49 ($B$) & F0V & $B$ & 2 & 6 & 2.2 & 0.78-0.81, 0.26-0.29 & -- & -- & A2 \\ And V342 & 7.98 ($B$) & A3 & $B$ & 1 & 6.5 & 1.8 & 0.90-1.00 & -- & -- & A3 \\ And V363 & 9.29 ($B$) & A2 & $B$ & 1 & 4 & 4.2 & 0.00-0.14 & -- & -- & A3 \\ Aql QY & 11.4 ($B$) & F0 & $BVI$ & 36 & 200+ & 3.8 & 0.00-1.00 &10.655 (3)& 12.0 (1)& A2 \& K \\ Aql V805 & 7.85 ($B$) & A3 & $B$ & 1 & 4 & 1.3 & 0.82-0.88 & -- & -- & A3 \\ Aql V1461 & 9.00 ($B$) & A0 & $B$ & 1 & 4.5 & 3.7 & 0.56-0.67 & -- & -- & A3 \\ Aqr CZ & 11.20 ($B$) & A5 & $B$ & 10 & 25+ & 2.9 & 0.00-1.00 &35.508 (2)& 3.7 (5) & A2 \\ Ari SZ & 11.60 ($B$) & F0 & $V$ & 1 & 3.5 & 3.6 & 0.51-0.60 & -- & -- & A2 \\ Aur V417 & 7.99 ($B$) & A0& $BVRI$ & 9 & 45+ & 1.5 & 0.00-1.00 & -- & -- & A2 \\ Cam SS & 10.93($B$) & G1III& $B$ & 1 & 4 & 3.8 & 0.61-0.64 & -- & -- & A2 \\ Cas IS & 12.10($B$) & A2& $B$ & 1 & 4 & 4.1 & 0.19-0.26 & -- & -- & K \\ Cas V364 & 11.10($B$) & A7& $B$ & 1 & 4 & 3.4 & 0.15-0.26 & -- & -- & A2 \\ Cas V773 & 6.32 ($B$) & A3& $B$ & 1 & 4 & 2.1 & 0.45-0.59 & -- & -- & A3 \\ Cas V821 & 8.37 ($B$) & A0& $B$ & 1 & 4 & 1.5 & 0.60-0.68 & ? & -- & A2 \\ Cep EI & 7.94 ($B$) & A5& $B$ & 1 & 4.5 & 1.4 & 0.22-0.24 & ? & -- & A2 \\ Cep V405 & 8.95 ($B$) & A2& $BVRI$ & 5 & 35 & 2.4 & 0.00-1.00 & -- & -- & A2 \\ Cep WX & 9.38 ($B$) & A3& $B$ & 1 & 2.5 & 3.8 & 0.64-0.67 & -- & -- & A2 \\ Cep XX & 9.47 ($B$) & A7V& $B$ & 1 & 4 & 2.5 & 0.63-0.69 & 32.07 (4)& 3.6 (4) & A2 \\ Cet DP & 7.01 ($B$) & A2& $B$ & 1 & 4 & 1.9 & 0.88-0.92 & -- & -- & A3 \\ Cyg MY & 8.68 ($B$) & A2.5& $B$ & 3 & 13 & 2.0 & 0.21-0.29, 0.71-0.76 & -- & -- & A2 \\ Cyg UW & 11.00 ($B$) & A5& $BVI$ & 26 & 90+ & 2.7 & 0.00-1.00 &27.841 (2)& 1.9 (2) & A2 \\ Cyg V477 & 8.71 ($B$) & A1V& $BV$ & 1 & 4 & 2.9 & 0.71-0.77 & -- & -- & A2 \\ Cyg V959 & 11.50 ($B$) & A5& $B$ & 1 & 4 & 3.4 & 0.49-0.61 & -- & -- & K \\ Cyg V2083 & 7.13 ($B$) & A3& $B$ & 1 & 4 & 2.3 & 0.09-0.15 & -- & -- & A3 \\ Cyg V2154 & 8.18 ($B$) & F0& $B$ & 1 & 4 & 1.1 & 0.59-0.65 & -- & -- & A2 \\ Cyg VW & 10.58 ($B$) & A3 & $B$ & 1 & 4 & 3.0 & 0.20-0.22 & -- & -- & A2 \\ Dra HL & 7.52 ($B$) & A5& $BVRI$ & 17 & 60+ & 2.1 & 0.00-1.00 &26.914 (1)& 3.0 (2) & A2 \\ Dra HZ & 8.34 ($B$) & A0& $BVRI$ & 8 & 25+ & 2.7 & 0.00-1.00 &51.068 (2)& 4.0 (4) & A3 \\ Dra RX & 10.84 ($B$) & F0& $B$ & 2 & 11 & 5.4 & 0.38-0.45, 0.58-0.64 & -- & -- & A2 \\ Her AD & 10.02 ($B$) & A4V& $B$ & 1 & 4 & 4.2 & 0.63-0.65 & -- & -- & A3 \\ Her BO & 11.6 ($B$) & A7 & $BVI$ & 26 & 130 & 4.5 & 0.00-1.00 &13.430 (1)& 68 (3) & A2 \& K \\ Her FN & 10.50 ($B$) & A8& $B$ & 1 & 4.5 & 6.0 & 0.76-0.83 & -- & -- & A2 \\ Her HS & 8.58 ($B$) & B6III & $BVI$ & 4 & 4 & 3.8 & 0.20-0.30 & -- & -- & A2 \\ Her SZ & 10.28 ($B$) & F0V& $B$ & 1 & 4 & 2.8 & 0.20-0.29 & -- & -- & A2 \\ Her UX & 9.11 ($B$) & A0V& $B$ & 1 & 4 & 3.6 & 0.73-0.82 & -- & -- & A1 \\ Her V948 & 9.26 ($B$) & F2& $BVRI$ & 10 & 35 & 3.5 & 0.00-1.00 & -- & -- & A2 \\ Her V1002 & 9.14 ($B$) & A0& $B$ & 1 & 4 & 3.5 & 0.13-0.23 & -- & -- & A3 \\ Hya DE & 11.00 ($B$) & A2& $B$ & 1 & 4 & 3.9 & 0.28-0.32 & -- & -- & A2 \\ Lac AU & 11.50 ($B$) & A5& $BVRI$ & 19 & 80+ & 2.0 & 0.00-1.00 &58.217 (1)& 5.0 (3) & K \& A2 \\ Lac CM & 8.39 ($B$) & A3& $BVRI$ & 10 & 40+ & 2.6 & 0.00-1.00 & ? & -- & A2 \\ Lac VX & 10.83 ($B$) & F0& $B$ & 3 & 14 & 4.5 & 0.10-0.50 & -- & -- & A2 \\ Lac V364 & 8.54 ($B$) & A3& $B$ & 1 & 4 & 1.9 & 0.25-0.27 & -- & -- & A3 \\ Lac V398 & 8.87 ($B$) & A0& $B$ & 1 & 4 & 3.1 & 0.27-0.30 & -- & -- & A3 \\ Lyn CL & 10.05 ($B$) & A5& $BVI$ & 12 & 90+ & 4.0 & 0.00-1.00 &23.051 (1)& 7.3 (3) & A4 \\ Lyn SX & 10.00 ($B$) & A2& $BVI$ & 6 & 30 & 8.4 & 0.00-0.10, 0.48-0.65 & -- & -- & A3 \\ Lyr RV & 11.50 ($B$) & A5& $B$ & 1 & 6 & 3.6 & 0.74-0.80 & -- & -- & K \\ Mon EP & 10.50 ($V$) & A3& $B$ & 1 & 4 & 3.6 & 0.20-0.34 & -- & -- & A2 \\ Mon HO & 11.40 ($B$) & A5& $B$ & 1 & 5.5 & 3.9 & 0.18-0.21 & -- & -- & A2 \\ Oph V391 & 11.50 ($B$) & A1& $B$ & 1 & 5 & 4.5 & 0.56-0.65 & -- & -- & K \\ Oph V456 & 10.37 ($B$) & A2& $B$ & 1 & 4 & 3.7 & 0.67-0.85 & -- & -- & A1 \\ Ori EY & 10.21 ($B$) & A7& $B$ & 1 & 4 & 7.8 & 0.32-0.33 & -- & -- & A4 \\ Ori FK & 11.80 ($B$) & A2& $B$ & 1 & 4 & 2.0 & 0.35-0.44 & -- & -- & K \\ Ori FT & 9.35 ($B$) & A0& $B$ & 1 & 4 & 2.7 & 0.72-0.77 & -- & -- & A2 \\ Ori V536 & 10.50 ($B$) & A2& $B$ & 1 & 3 & 3.1 & 0.46-0.50 & -- & -- & A1 \\ Peg AT & 9.21 ($B$) & A3.5V& $BR$ & 8 & 30 & 3.3 & 0.00-1.00 & -- & -- & A3 \\ Peg BG & 10.50 ($B$) & A2V & $BVRI$ & 15 & 102 & 5.8 & 0.00-0.29, 0.35-1.00 &25.543 (1)& 12.7 (5) & A2 \& A4 \\ Peg DM & 11.80 ($B$) & A3& $B$ & 1 & 4 & 4.3 & 0.72-0.79 & ? & -- & A2 \\ Peg OO & 8.53 ($B$) & A2& $B$ & 1 & 4 & 2.9 & 0.79-0.83 & -- & -- & A3 \\ Per RV & 11.40 ($B$) & A0& $B$ & 1 & 5 & 3.9 & 0.16-0.27 & -- & -- & A2 \\ Sge UZ & 11.40 ($B$) & A0& $VR$ & 3 & 60+ & 4.9 & 0.00-1.00 & ? & -- & A2 \\ Tau EW & 11.70 ($B$) & - & $B$ & 1 & 4 & 3.1 & 0.59-0.62 & -- & -- & K \\ Tau V1149 & 8.65 ($B$) & A0& $B$ & 1 & 4.5 & 2.1 & 0.63-0.67 & -- & -- & A4 \\ UMa IO & 8.44 ($B$) & A3& $BVRI$ & 47 & 150+ & 1.5 & 0.00-1.00 &22.015 (2)& 6.7 (1) & A2 \& A3 \\ UMi RT & 11.10 ($B$) & F0& $B$ & 1 & 5 & 3.7 & 0.49-0.61 & -- & -- & A2 \\ Vul AW & 10.00 ($B$) & F0& $B$ & 1 & 5.5 & 4.4 & 0.59-0.88 & -- & -- & A2 \\ Vul BP & 10.17 ($B$) & A7& $B$ & 1 & 3 & 2.5 & 0.24-0.30 & -- & -- & A2 \\ Vul RR & 10.15 ($B$) & A2& $B$ & 2 & 6 & 2.7 & 0.16-0.18, 0.64-0.67 & ? & -- & A2 \\ \hline \end{tabular}} \end{table*} \section[]{Spectroscopic analysis} WY~Cet, UW~Cyg and a total of 23 spectroscopic standard stars ranging from A0 to G8 spectral types were observed with the same instrumental set-up. We used 900-s and 2000-s exposure times for WY~Cet and UW~Cyg, respectively. Both spectra were observed well inside the secondary eclipses (at the phase 0.506 for WY~Cet and 0.509 for UW~Cyg) when the light contribution from the secondary component is minimal, therefore the spectra practically correspond to the primaries. All spectra were calibrated and normalized to enable direct comparisons. Then, we shifted the spectra, using H$_{\beta}$ as reference, to compensate for the relative Doppler shifts of each standard. The variables' spectra were subtracted from those of each standard, deriving sums of squared residuals in each case. Such least squares sums should allow the best match between the spectra of variable and standard to be found. \begin{table} \centering \caption{The photometric observations log of the eight selected EBs.} \begin{tabular}{lll} \hline System & Comparison star & Check star \\ \hline CZ Aqr & GSC 6396-1024 & GSC 6396-0872 \\ TY Cap & GSC 5749-2167 & GSC 5749-1557 \\ WY Cet & HIP 7373 & GSC 5279-0617 \\ UW Cyg & GSC 3164-0083 & GSC 3164-0269 \\ HL Dra & SAO 31053 & GSC 3913-0901 \\ HZ Dra & SAO 18500 & GSC 4449-1053 \\ AU Lac & GSC 3610-0231 & GSC 3610-0685 \\ CL Lyn & GSC 3787-0420 & GSC 3783-0649 \\ \hline \end{tabular} \end{table} Comparison between the spectra of WY~Cet and UW~Cyg and the standards yielded the primaries to be A9V and A6V type stars, respectively. Fig.~1a shows the best matching of the variables' spectra with those of the standards. The spectrum of HZ~Dra, although dominated by the primary component, could not fit sufficiently well with any of the standards. That may means that the secondary contributes in way that affects the spectrum, therefore spectral classification could not be accurate. For the radial velocity (RV) calculations for HZ~Dra the software \emph{Broadening Functions} (BFs) v.2.4c \citep{NE09}, which is based on the method of \citet{RU02}, was used. We cropped all spectra in order to avoid the broad H$_{\beta}$ line, and we included all the sharp metallic lines between 4800-5350~\AA. Each RV value and its error was derived statistically (mean value and error) from the respective velocities resulting from BFs method by using six different standard stars of similar spectral type to the system. Due to the large brightness difference between the components of the system, we obtained measurements only for its primary. The semi-amplitude of the RV curve, $K_1$, and the systemic velocity $V_0$ were calculated by fitting a sinusoidal function to the RV points. A sample of the heliocentric RVs are given in Table~4, while the rest are given in the electronic version of the paper. The RV plot is illustrated in Fig.~1b. \begin{table} \centering \caption{Sample of the heliocentric radial velocities measurements of HZ~Dra.} \begin{tabular}{lcc} \hline HJD & Phase & $RV_1$ \\ & & (km/s) \\ \hline 2455825.3099 & 0.2800 & $-$36 (14) \\ 2455825.3549 & 0.3382 & $-$35 (14) \\ 2455829.4886 & 0.6862 & 28 (12) \\ \hline \end{tabular} \end{table} \begin{figure} \begin{tabular}{cc} \includegraphics[width=7.2cm]{spectra.eps}&(a)\\ \includegraphics[width=7.2cm]{Dra_HZ_RV.eps}&(b)\\ \end{tabular} \caption{(a): The comparison spectra of WY~Cet (upper) and UW~Cyg (lower) and the standard stars A9V (HIP~11678) and A6V (HIP~21589), respectively. (b): Synthetic (solid line) and observed (points) radial velocities of the primary component of HZ~Dra. The radial velocity amplitude $K_1$ and the systemic velocity $V_0$ are also indicated.} \label{fig1} \end{figure} \section[]{Light curve analysis and absolute parameters derivation} Complete LCs of each system were analysed using \emph{PHOEBE} v.0.29d software \citep{PZ05} that follows the 2003 version of the Wilson-Devinney (WD) code \citep{WD71,WI79,WI90}. In the cases where multicolour photometry was available, the LCs were analysed simultaneously. For HL~Dra and HZ~Dra the radial velocities of the primary component (\citet{PR06} and present paper) were included in the analysis. In the absence of spectroscopic mass ratios, the `$q$-search' method using a step of 0.1 was trialled in Modes 2 (detached system) and 5 (conventional semi-detached system) to find `photometric' estimates for the mass ratio $q_{\rm ph}$. This value was then set as initial input and treated as a free parameter in the subsequent analysis. The temperatures of the primaries were assigned values according to their spectral types using the correlations of \citet{CO00} and were kept fixed, while the temperatures of the secondaries $T_2$ were adjusted. The values of bolometric albedos $A_1$ and $A_2$, and gravity darkening coefficients, $g_1$ and $g_2$, were set as $A$=1 and $g$=1 for radiative \citep{RU69,VZ24} and $A$=0.5 and $g$=0.32 for convective atmospheres \citep{RU69,LU67}. Synchronous rotation was assumed, so the synchronization parameters $F_1$ and $F_2$ were set as 1. The linear limb darkening coefficients, $x_1$ and $x_2$, were taken from the tables of \citet{VH93}; the dimensionless potentials $\Omega_{1}$ and $\Omega_{2}$, the fractional luminosity of the primary component $L_{1}$ and the inclination $i$ of the system's orbit were set as adjustable. Since the O$-$C diagrams (see next section) of CZ~Aqr, TY~Cap, WY~Cet and UW~Cyg suggested possible existence of other components, the third light parameter $l_3$ was also adjusted. Best-fit models and observed LCs of the systems are presented in Fig.~2 with corresponding parameters in Table~5. \begin{landscape} \begin{table} \centering \caption{Light curve solution and absolute parameters of the components (P=Primary, S=Secondary).} \scalebox{0.85}{ \begin{tabular}{l cccc cccc cccc cccc} \hline System:&\multicolumn{2}{c}{CZ Aqr} & \multicolumn{2}{c}{TY Cap}&\multicolumn{2}{c}{WY Cet} & \multicolumn{2}{c}{UW Cyg}&\multicolumn{2}{c}{HL Dra} & \multicolumn{2}{c}{HZ Dra}&\multicolumn{2}{c}{AU Lac} & \multicolumn{2}{c}{CL Lyn}\\ \hline \multicolumn{16}{c}{\textsl{Light curve parameters}}\\ \hline Mode & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Detached} & \multicolumn{2}{c}{Semidetached} & \multicolumn{2}{c}{Semidetached} \\ $i~(\degr$) & \multicolumn{2}{c}{89.7 (1)} & \multicolumn{2}{c}{80.4 (2)} & \multicolumn{2}{c}{81.8 (1)} & \multicolumn{2}{c}{87.1 (1)} & \multicolumn{2}{c}{66.5 (1)} & \multicolumn{2}{c}{72.0 (3)} & \multicolumn{2}{c}{83.0 (1)} & \multicolumn{2}{c}{78.7 (1)} \\ $q~(m_{2}/m_{1}$) & \multicolumn{2}{c}{0.49 (1)} & \multicolumn{2}{c}{0.52 (1)} & \multicolumn{2}{c}{0.26 (1)} & \multicolumn{2}{c}{0.14 (1)} & \multicolumn{2}{c}{0.37 (1)} & \multicolumn{2}{c}{0.12 (4)} & \multicolumn{2}{c}{0.30 (1)} & \multicolumn{2}{c}{0.19 (2)} \\ \hline \textsl{Component:}& P & S & P & S & P & S & P & S & P & S & P & S & P & S & P & S \\ \hline $T$ (K) & 8200$^1$ & 5650 (12) & 8200$^1$ & 4194 (30) & 7500$^2$ & 4347 (7) & 8000$^2$ & 4347 (4) & 8200$^3$ & 5074 (8) & 9800$^4$ & 5015 (68) & 8200$^5$ & 3784 (15) & 8200$^6$ & 4948 (14) \\ $\Omega$ & 3.44 (1) & 2.86 & 3.80 (2) & 2.93 & 4.18 (1) & 2.40 & 5.89 (1) & 2.07 & 2.93 (1) & 2.63 & 2.48 (1) & 2.28 (2) & 4.51 (2) & 2.46 & 3.36 (1) & 2.21 \\ $x_{\rm B}$ & 0.584 & 0.760 & 0.614 & 1.004 & 0.701 & 0.968 & 0.596 & 0.971 & 0.596 & 0.853 & 0.491 & 0.862 & 0.540 & 0.831 & 0.588 & 0.874 \\ $x_{\rm V}$ & -- & -- & 0.510 & 0.850 & 0.570 & 0.822 & 0.517 & 0.820 & 0.509 & 0.708 & 0.418 & 0.721 & 0.474 & 0.747 & 0.505 & 0.729 \\ $x_{\rm R}$ & -- & -- & 0.424 & 0.738 & 0.476 & 0.708 & -- & -- & 0.426 & 0.611 & 0.353 & 0.622 & 0.402 & 0.691 & -- & -- \\ $x_{\rm I}$ & -- & -- & 0.331 & 0.606 & 0.367 & 0.586 & 0.351 & 0.591 & 0.338 & 0.513 & 0.281 & 0.523 & 0.327 & 0.564 & 0.340 & 0.527 \\ ($L/L_{\rm T})_{\rm B}$ & 0.878 (2) & 0.122 (1) & 0.953 (4) & 0.018 (1) & 0.946 (2) & 0.044 (1) & 0.945 (2) & 0.051 (1) & 0.956 (1) & 0.044 (1) & 0.995 (1) & 0.005 (1) & 0.987 (1) & 0.013 (1) & 0.959 (6) & 0.042 (2) \\ ($L/L_{\rm T})_{\rm V}$ & -- & -- & 0.938 (6) & 0.037 (1) & 0.915 (2) & 0.076 (1) & 0.902 (3) & 0.093 (2) & 0.932 (1) & 0.068 (1) & 0.992 (1) & 0.008 (1) & 0.969 (1) & 0.031 (1) & 0.934 (8) & 0.066 (4) \\ ($L/L_{\rm T})_{\rm R}$ & -- & -- & 0.914 (8) & 0.056 (1) & 0.883 (2) & 0.106 (2) & -- & -- & 0.912 (1) & 0.088 (1) & 0.988 (1) & 0.012 (2) & 0.948 (2) & 0.052 (2) & -- & -- \\ ($L/L_{\rm T})_{\rm I}$ & -- & -- & 0.887 (10) & 0.084 (2) & 0.843 (3) & 0.144 (3) & 0.802 (4) & 0.180 (1) & 0.888 (1) & 0.112 (1) & 0.984 (1) & 0.016 (4) & 0.916 (2) & 0.084 (2) & 0.887 (10) & 0.113 (1) \\ ($L_3/L_{\rm T})_{\rm B}$ & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{0.029 (3)} & \multicolumn{2}{c}{0.010 (1)} & \multicolumn{2}{c}{0.004 (1)} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} \\ ($L_3/L_{\rm T})_{\rm V}$ & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{0.025 (4)} & \multicolumn{2}{c}{0.009 (2)} & \multicolumn{2}{c}{0.005 (1)} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} \\ ($L_3/L_{\rm T})_{\rm R}$ & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{0.030 (6)} & \multicolumn{2}{c}{0.010 (2)} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} \\ ($L_3/L_{\rm T})_{\rm I}$ & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{0.029 (8)} & \multicolumn{2}{c}{0.013 (2)} & \multicolumn{2}{c}{0.018 (1)} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} & \multicolumn{2}{c}{--} \\ \hline \multicolumn{16}{c}{\textsl{Absolute parameters}}\\ \hline $M~(M_{\sun}$) & 2.00$^a$ & 0.98 (1) & 2.00$^a$ & 1.05 (1) & 1.70$^a$ & 0.44 (1) & 1.90$^a$ & 0.26 (1) & 2.5 (2) & 0.9 (1) & 3.0 (3) & 0.4 (1) & 2.0$^a$ & 0.60 (1) & 2.0$^a$ & 0.38 (3) \\ $R~(R_{\sun}$) & 1.9 (1) & 1.8 (1) & 2.5 (1) & 2.5 (1) & 2.2 (1) & 2.3 (1) & 2.2 (1) & 2.9 (1) & 2.5 (4) & 1.8 (3) & 2.3 (1) & 0.8 (1) & 1.8 (1) & 2.1 (1) & 2.5 (1) & 1.9 (1) \\ $L~(L_{\sun}$) & 15.3 (9) & 2.9 (2) & 24.3 (8) & 1.8 (2) & 14.0 (9) & 1.7 (1) & 18.0 (9) & 2.6 (1) & 24.3 (7) & 1.9 (1) & 45 (3) & 0.4 (2) & 12.6 (7) & 0.8 (1) & 25.2 (9) & 2.0 (7) \\ $M_{\rm bol}$ (mag) & 1.8 (6) & 3.6 (6) & 1.3 (1) & 4.1 (1) & 1.9 (8) & 4.2 (8) & 1.6 (4) & 3.7 (6) & 1.3 (2) & 4.1 (2) & 0.6 (4) & 5.9 (4) & 2.0 (6) & 5.0 (7) & 1.2 (9) & 4.0 (8) \\ \textsl{a}~($R_{\sun}$) & 1.9 (2) & 3.8 (1) & 2.7 (3) & 5.2 (1) & 1.8 (3) & 6.9 (1) & 1.5 (2) & 11.2 (1) & 1.7 (3) & 4.4 (1) & 0.6 (1) & 4.7 (2) & 1.7 (2) & 5.7 (1) & 1.3 (3) & 6.6 (1) \\ \hline \multicolumn{17}{l}{$^1$\citet{PE97}, $^2$present paper (section~3), $^3$\citet{PR06}, $^4$\citet{WR03}, $^5$\citet{BU04}, $^6$\citet{AD01}, $^a$assumed, $L_{\rm T}=L_1+L_2+L_3$} \end{tabular}} \end{table} \begin{figure} \begin{tabular}{cccc} \includegraphics[width=5.7cm]{Aqr_CZ_LC.eps}&\includegraphics[width=5.7cm]{Cap_TY_LC.eps}&\includegraphics[width=5.7cm]{Cet_WY_LC.eps}&\includegraphics[width=5.7cm]{Cyg_UW_LC.eps}\\ \includegraphics[width=5.7cm]{Dra_HL_LC.eps}&\includegraphics[width=5.7cm]{Dra_HZ_LC.eps}&\includegraphics[width=5.7cm]{Lac_AU_LC.eps}&\includegraphics[width=5.7cm]{Lyn_CL_LC.eps}\\ \end{tabular} \caption{Synthetic (solid lines) and observed (grey points) light curves of the systems.} \label{fig2} \end{figure} \end{landscape} Although no radial velocity curves for these systems were published as yet, except for the primaries of HL~Dra and HZ~Dra, we can form fair estimates of their absolute parameters (see Table~5). The masses of the primaries were assumed from their spectral type, while those of the secondaries followed from the adopted mass ratios. The semi-major axes (\textsl{a}), used to calculate mean radii, follow from Kepler's law. In Fig.~3 the position of the systems' components in the Mass-Radius ($M-R$) diagram is presented. The theoretical lines for Zero Age Main Sequence (ZAMS) and Terminal Age Main Sequence (TAMS) were taken from \citet{NM03}. \begin{figure} \includegraphics[width=8cm]{M-R.eps} \caption{The location of the systems' components ($P$ for primary and $S$ for the secondary) in the $M-R$ diagram. The primaries and the secondaries belonging to near contact systems are indicated with grey symbols for comparison.} \label{fig3} \end{figure} | 14 | 4 | 1404.2808 |
|
1404 | 1404.7100_arXiv.txt | {The $f(T,T_G)$ class of gravitational modification, based on the quadratic torsion scalar $T$, as well as on the new quartic torsion scalar $T_G$ which is the teleparallel equivalent of the Gauss-Bonnet term, is a novel theory, different from both $f(T)$ and $f(R,G)$ ones. We perform a detailed dynamical analysis of a spatially flat universe governed by the simplest non-trivial model of $f(T,T_G)$ gravity which does not introduce a new mass scale. We find that the universe can result in dark-energy dominated, quintessence-like, cosmological-constant-like or phantom-like solutions, according to the parameter choices. Additionally, it may result to a dark energy - dark matter scaling solution, and thus it can alleviate the coincidence problem. Finally, the analysis ``at infinity'' reveals that the universe may exhibit future, past, or intermediate singularities depending on the parameters.} | \label{Introduction} Since the discovery of the universe late-times acceleration, a large amount of research has been devoted to its explanation. In principle, one can follow two main directions to achieve this. The first way is to modify the content of the universe introducing the dark energy concept, with its simpler candidates being a canonical scalar field, a phantom field or the combination of both fields in a unified model dubbed quintom (for reviews on dark energy see \cite{Copeland:2006wr,Cai:2009zp} and references therein). The second direction that one can follow is to modify the gravitational sector itself (for a review see \cite{Capozziello:2011et} and references therein), acquiring a modified cosmological dynamics. However, note that apart from the interpretation, one can transform from one approach to the other, since the crucial issue is just the number of degrees of freedom beyond General Relativity and standard model particles (see \cite{Sahni:2006pa} for a review on such a unified point of view). Finally, note that the above scenarios, apart from late-times implications, can be also used for the description of the inflationary stage \cite{Nojiri:2003ft}. In the majority of modified gravitational theories, one suitably extends the curvature-based Einstein-Hilbert action of General Relativity. However, an interesting class of gravitational modification arises when one modifies the action of the equivalent formulation of General Relativity based on torsion. In particular, it is known that Einstein himself constructed the so-called ``Teleparallel Equivalent of General Relativity'' (TEGR) \cite{Unzicker:2005in,Hayashi:1979qx,Pereira,Maluf:2013gaa} using the curvature-less Weitzenb{\"{o}}ck connection instead of the torsion-less Levi-Civita one. The corresponding Lagrangian, namely the torsion scalar $T$, is constructed by contractions of the torsion tensor, in a similar way that the usual Einstein-Hilbert Lagrangian $R$ is constructed by contractions of the curvature (Riemann) tensor. Thus, inspired by the $f(R)$ modifications of the Einstein-Hilbert Lagrangian \cite{DeFelice:2010aj,Nojiri:2010wj}, one can construct the $f(T)$ modified gravity by extending $T$ to an arbitrary function \cite{Ferraro:2006jd,Ben09,Linder:2010py}. Note that although TEGR coincides with General Relativity at the level of equations, $f(T)$ does not coincide with $f(R)$, that is they represent different modification classes. Thus, the cosmological implications of $f(T)$ gravity are new and interesting \cite{Linder:2010py,Chen:2010va,Dent:2011zz,Zheng:2010am, Sharif001,Li:2011rn, Cai:2011tc,Boehmer:2011gw,Capozziello:2011hj, Daouda:2011rt,Geng:2011aj,Wu:2011kh,Gonzalez:2011dr,Wei:2011aa, Atazadeh:2011aa, Farajollahi:2011af,Karami:2012fu,Iorio:2012cm,Cardone:2012xq, Capozziello:2012zj,Jamil:2012ti, Ong:2013qja, Amoros:2013nxa, Otalora:2013dsa, Geng:2013uga,Nesseris:2013jea, Bamba:2013ooa, Nashed:2014uta,Harko:2014sja,Harko:2014aja}. However, in curvature gravity, apart from the simple $f(R)$ modification one can construct more complicated extensions using higher-curvature corrections such as the Gauss-Bonnet term $G$ \cite{Wheeler:1985nh,Antoniadis:1993jc,Nojiri:2005vv} or functions of it \cite{Nojiri:2005jg,DeFelice:2008wz}, Lovelock combinations \cite{Lovelock:1971yv,Deruelle:1989fj,Charmousis:2008kc}, and Weyl combinations \cite{Mannheim:1988dj,Flanagan:2006ra,Grumiller:2013mxa}. Inspired by these, in the recent work \cite{Kofinas:2014owa}, the $f(T,T_G)$ gravitational modification was constructed, which is based on the old quadratic torsion scalar $T$, as well as on the new quartic torsion scalar $T_G$ that is the teleparallel equivalent of the Gauss-Bonnet term. Obviously, $f(T,T_G)$ theories cannot arise from the $f(T)$ ones, and additionally they are different from $f(R,G)$ class of curvature modified gravity. Thus, $f(T,T_G)$ is a novel class of gravitational modification. The cosmological applications of $f(T,T_G)$ gravity proves to be very interesting \cite{Kofinas:2014owa}. Therefore, it is both interesting and necessary to perform a dynamical analysis, examining in a systematic way the allowed cosmological behaviors, focusing on the late-times stable solutions. The phase-space and stability analysis is a very powerful tool, since it reveals the global features of a given cosmological scenario, independently of the initial conditions and the specific evolution of the universe. In the present investigation we perform such a detailed phase-space analysis, and we extract the late-times, asymptotic solutions, calculating also the corresponding observable quantities, such as the deceleration parameter, the effective dark energy equation-of-state parameter, and the various density parameters. The plan of the work is the following: In section \ref{fTgravity} we briefly review the scenario of $f(T,T_G)$ gravity and in section \ref{fTcosmology} we present its application in cosmology. In section \ref{Phasespaceanalysis} we perform the detailed dynamical analysis for the simplest non-trivial model of $f(T,T_{G})$ gravity. In section \ref{implications} we discuss the cosmological implications and the physical behavior of the scenario. Finally, in section \ref{Conclusions} we summarize our results. | \label{Conclusions} In the present work we studied the dynamical behavior of the recently proposed scenario of $f(T,T_G)$ cosmology \cite{Kofinas:2014owa}. This class of modified gravity is based on the quadratic torsion scalar $T$, which is the Lagrangian of the teleparallel equivalent of General Relativity, as well as on the new quartic torsion scalar $T_G$, which is the teleparallel equivalent of the Gauss-Bonnet term. Obviously, $f(T,T_G)$ theories are more general and cannot be spanned by the simple $f(T)$ ones, and additionally they are different from $f(R,G)$ class of curvature modified gravity too. Without loss of generality, as a simple, but non-trivial example, capable of revealing the advantages and the new features of the theory, we considered a model where $T$ and $T_G$ corrections are of the same order, and thus expected to play an important role at late times. We performed for a spatially flat universe the complete and detailed phase-space behavior, both in the finite and infinite regions, calculating additionally also the values of basic observables such is the various density parameters, the deceleration parameter and the dark energy equation-of-state parameter. This scenario exhibits interesting cosmological behaviors. In particular, depending on the model parameters, the universe can result in a dark energy dominated accelerating solution and the dark energy equation-of-state parameter can lie in the quintessence regime, it can be equal to the cosmological constant value $-1$, or it can even lie in the phantom regime. Additionally, it can result in a dark energy - dark matter scaling solution, and thus it can alleviate the coincidence problem. Finally, under certain parameter choices the universe can result to Big Rip, sudden, or other form of singularities, as it is usual in many modified gravitational theories. Definitely, before the scenario at hand can be considered as a good candidate for the description of Nature, a detailed confrontation with observations should be performed. In particular, one should use data from local gravity experiments (Solar System observations), as well as type Ia Supernovae (SNIa), Baryon Acoustic Oscillations (BAO), and Cosmic Microwave Background (CMB) radiation data, in order to impose constraints on the model. These necessary investigations lie beyond the scope of the present work and are left for a future project. | 14 | 4 | 1404.7100 |
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