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1404 | 1404.6108_arXiv.txt | {In this paper we model the vertical structure of mass distribution of the Milky Way galaxy in the framework of a finite-width global disk model. Assuming the Galactic rotation curve only, we test inside the solar orbit the predictions of the model for two measurable and unrelated to each other processes: the gravitational microlensing that allows to fix the disk width-scale by the best fit to measurements, and the vertical gradient of rotation modelled in the quasi-circular orbits approximation. The former is sensitive to the gravitating mass in compact objects and the latter is sensitive to all kinds of gravitating matter. The analysis points to a small width-scale of the considered disks \new{and, at most, insignificant contribution of non-baryonic dark mater in the solar circle. The predicted high vertical gradient values in the rotation are consistent with the gradient measurements.}} | The Milky Way (\mw) is an example of a galaxy with high vertical gradients of rotation measured at low altitudes above the mid-plane \citep{2008ApJ...679.1288L}. It is interesting to see the effect of the disk thickness on the gradient value prediction and to compare it with the analogous prediction in an infinitesimally thin global disk model \citep{2010MNRAS.407.1689J}. The gravitational microlensing phenomenon provides another constraint on the mass distribution, independent of the vertical gradient structure. In particular, the amount of mass seen through the gravitational microlensing measurements inside the solar orbit, was shown to be consistent with the dynamical mass ascertained from the Galactic rotation and reduced by the gas contribution undetectable by the microlensing \citep{2012A&A...546A.126S}. In the present paper we use microlensing measurements to constrain the disk width and we compare the resulting vertical gradient predictions with the gradient measurements. In our several previous studies, e.g., \citet{2010MNRAS.407.1689J,2011MNRAS.412..331J}, we modelled spiral galaxies in the approximation of an infinitesimally thin disk. In this framework, we could easily obtain high values of the vertical gradient of rotation, in accord with gradient measurements. However, the model does not account for the vertical structure in the mass distribution in the direct neighborhood of the mid-plane (the $z=0$ vicinity). This structure can be modelled by considering a finite-width disk \new{with an assumed vertical profile of mass density}. A preliminary step towards this we presented in \citet{2012A&A...546A.126S}, where a column mass density \new{of a disk with the exponential vertical profile} was identified with the surface density of an infinitesimally thin disk. \add{But this was a simple de-projection which did not take into account the redistribution of mass required to preserve the shape of rotation curve in the Galactic center vicinity.} In the next approximation considered in this paper, in \secref{sec:DiskTheory} and \secref{sec:lensing1}, we find \new{by iterations the volume mass densities of two example finite-width disks}. These disks exactly account for the \new{\mw's tabulated} rotation curve \new{and have their width-scales constrained by the gravitational microlensing measurements. In \secref{sec:gradient} we compare the predictions of the two models with the measurements of the vertical gradient of rotation}. | \add{ We obtained mass distribution in finite-width disks in an iterative fashion from a given rotation curve. We assumed that the volume density $\rho(r,z)$ could be factorized as $\rho(r,z)=\sigma(r)f(z)$, where $f(z)$ is a normalized vertical profile, with a characteristic width-scale which is a free parameter. We did not assume any constraints on the column mass density $\sigma(r)$, so that its functional dependence was governed entirely by the shape of the rotation curve. \new{The density effectively describes all forms of the dynamical mass inferred from the Galactic rotation, therefore, the disk-width should not be confused with the width-scale parameters measured in the Sun vicinity for various stellar subsystems.} \footnote{\new{Incidentally, the obtained width-scale of $\approx100\pc$ is consistent with the distribution of \newII{massive stars} in the Sun vicinity. \newII{However, massive} \newII{stars are rare locally and do not contribute substantially to the mass for any proposed IMF.}}} \new{Based on the gravitational microlensing measurements, it was possible within the disk model framework to determine the effective width-scales. We tested such obtained mass distributions by comparing the model predictions for the vertical gradient of the azimuthal velocity with the gradient measurements.} The (integrated) optical depth from the microlensing measurements is influenced by the amount of mass distributed along the lines of sight towards the Galactic center, whereas the details of the distribution are less important. We inferred the optimum width-scales of the considered disks by means of finding the best fits to the optical depth measured along various lines of sight. This shows that microlensing can be used as a tool to independently constrain the mass distribution models. With such determined width-scales, the resulting prediction for the behavior of the vertical gradient of rotation was compared with the gradient measurements in the mid-plane vicinity. This comparison turned out consistent with a small disk thickness. \new{(Interestingly enough, for such determined width-scales, the effective disk widths defined by the \Q{1/e criterion}, are almost equal: $2h'=176\pc$ and $\sim1.49h=174\pc$, respectively, for the exponential and for the \mexhat vertical profiles.)} The behavior of the vertical gradient of the azimuthal velocity and its value, when calculated on low altitudes above the mid-plane, is very sensitive to the width-scale parameter. At a given altitude in the gradient measurements region, the calculated gradient value changes significantly with the width-scale parameter. When the parameter is too high the absolute gradient value is too low compared with the measurements. Higher absolute gradients at low altitudes above the mid-plane suggest a smaller effective thickness of the Galactic disk. The rotation velocity is another factor that governs the gradient value. In particular, given a disk thickness and the gradient behavior, one could constrain the allowable range for the motion of the standard of rest at the Sun position (the width-scale could be increased for higher $\vsun$). Testing the vertical structure of the mass distribution with the help of the gradient measurements is thus a particularly sensitive tool, and this is therefore important to have high accuracy measurements of the gradient at small altitudes above the mid-plane. } Column mass density of flattened mass distributions is sensitive to uncertainties in the circular velocity. This sensitivity can be observed in the approximation of infinitely thin disk model \citep{1987gady.book.....B}. It is therefore important to have reliable rotation curves when studying flattened galaxies. However, Galactic rotation is relatively well known inside solar circle, therefore, with better data, we can expect some differences in the column density to occur close to or outside solar circle. These changes should not influence significantly results for global quantities inside solar circle, such as the predicted integrated optical depth \newIII{or vertical gradient inside the solar circle. In particular, as shown in \secref{app:toymodel}, with suitably changed external part of the rotation curve, the disk model prediction for the column density at the solar circle of $\approx140\,\msun\pc^{-2}$ made with the present rotation curve, could be reduced to a local value of $\approx70\,\msun\pc^{-2}$ inferred at the Sun vicinity from Jeans modelling \citep{1991ApJ...367L...9K,2013ApJ...772..108Z}. } | 14 | 4 | 1404.6108 |
1404 | 1404.0032_arXiv.txt | {% We present high-resolution, high dynamic range column-density and color-temperature maps of the Orion complex using a combination of \textit{Planck} dust-emission maps, \textit{Herschel} dust-emission maps, and 2MASS NIR dust-extinction maps. The column-density maps combine the robustness of the 2MASS NIR extinction maps with the resolution and coverage of the \textit{Herschel} and \textit{Planck} dust-emission maps and constitute the highest dynamic range column-density maps ever constructed for the entire Orion complex, covering $\SI{0.01}{mag} < A_K < \SI{30}{mag}$, or $\SI{2D20}{cm^{-2}} < N < \SI{5D23}{cm^{-2}}$. We determined the ratio of the \SI{2.2}{\um} extinction coefficient to the \SI{850}{\um} opacity and found that the values obtained for both Orion~A and B are significantly lower than the predictions of standard dust models, but agree with newer models that incorporate icy silicate-graphite conglomerates for the grain population. We show that the cloud projected pdf, over a large range of column densities, can be well fitted by a simple power law. Moreover, we considered the local Schmidt-law for star formation, and confirm earlier results, showing that the protostar surface density $\Sigma_*$ follows a simple law $\Sigma_* \propto \Sigma_\mathrm{gas}^\beta$, with $\beta \sim 2$.} | \label{sec:introduction} Our inability to accurately map the distribution of gas inside a molecular cloud has been a major impediment to understanding the star formation process. This is because tracing mass in molecular clouds is challenging when about 99\% of the mass of a cloud is in the form of H$_2$ and helium, which are invisible to direct observation at the cold temperatures that characterize these clouds. Tracing mass in molecular clouds is currently achieved through use of column-density tracers, such as molecular-line emission, thermal dust-emission, and dust-extinction. The simplest and most straightforward of these by far is dust-extinction, in particular, near-infrared (NIR) dust-extinction, as it directly traces the dust opacity (without assumptions on the dust temperature), and relies on the well-behaved optical properties of dust grains in the NIR (e.g., \citealp{2013A&A...549A.135A}). The advantages of the NIR dust-extinction technique as a column-density tracer have been discussed independently by \citet{2009ApJ...692...91G}, who performed an unbiased comparison between the three standard density-tracer methods, namely, NIR dust-extinction (\textsc{Nicer}, \citealp{2001A&A...377.1023L}), dust thermal emission in the millimeter and far-IR, and molecular-line emission. These authors found that dust-extinction is a more reliable column-density tracer than molecular gas (CO), and that observations of dust-extinction provide more robust measurements of column-density than observations of dust-emission (because of the dependence of the latter on the uncertain knowledge of dust temperatures, $T$, and dust emissivities, $\beta$). This implies that in a massive star-forming cloud, where cloud temperatures can vary significantly because of the large number of embedded young stars and protostars, dust-emission maps are fundamentally limited as tracers of cloud mass. This is particularly true for the densest cloud regions where star formation takes place. Although straightforward and robust, the NIR extinction technique is nevertheless limited by the number of available background stars that are detectable through a cloud \citep{1994ApJ...429..694L, 1998ApJ...506..292A, 2001A&A...377.1023L, 2009A&A...493..735L}. This implies that the resolution of a NIR extinction map is a function of Galactic latitude and, to a minor extent, Galactic longitude. For example, angular resolutions on the order of \SI{10}{arcsec} are easily achievable toward the Galactic Bulge with modern NIR cameras on \SI{10}{m} class telescopes (corresponding to a physical resolution of the order of \SI{1000}{AU} for regions such as the Pipe Nebula, Ophiuchus, Lupus, and Serpens). But for regions of critical importance for star formation, such as Orion~A, the cloud hosting the nearest massive star formation region to Earth, only angular resolutions of about \SI{1}{arcmin} (or physical resolutions on the order of \SI{24000}{AU}) are currently achievable with similar instrumentation, because of the location of this cloud toward the anti-center of the Galaxy, and about $20^\circ$ off the Galactic plane. The recent public release of ESA's \textit{Planck} and \textit{Herschel} thermal dust-emission data offers an excellent opportunity to study entire giant molecular complexes away from the Galactic plane, such as Orion, at resolutions on the order of 5000~AU, or about five times better than what is currently possible with NIR extinction techniques. The \textit{Planck} space observatory \citep{2010A&A...520A...1T,2011A&A...536A...1P} is an ESA space observatory launched on 14 May 2009 to measure the anisotropy of the cosmic microwave background (CMB). It observes the sky in nine frequency bands covering \SIrange{30}{857}{GHz} with high sensitivity and angular resolution from \SI{3}{arcmin} to \SI{5}{arcmin}. Most relevant to the study of thermal dust-emission from molecular clouds, the High Frequency Instrument (HFI; \cite{2010A&A...520A...9L, 2011A&A...536A...6P}) covers the \SIlist{100;143; 217;353;545;857}{GHz} (or \SIlist{3000;2100;1400;850;550;350}{\um} respectively) bands with bolometers cooled to \SI{0.1}{K}, providing a large-scale view of entire molecular complexes with an unprecedented sensitivity to dust-emission. The \textit{Herschel} space observatory \citep{2010A&A...518L...1P} is an ESA space observatory working in the far-infrared and submillimeter bands. The high sensitivity of \textit{Herschel} imaging cameras PACS \citep{2010A&A...518L...2P} and SPIRE \citep{2010A&A...518L...3G} are able to generate dust-emission maps with dynamic ranges that are not possible from ground-based bolometers, and reaching low column densities similar to those reached by NIR dust-extinction, although with a uniform resolution across the sky (of about \SI{12}{arcsec} at \SI{160}{\um}, \SI{18}{arcsec} at \SI{250}{\um}, and \SI{36}{arcsec} at \SI{500}{\um}). Unlike the \textit{Planck} satellite, however, \textit{Herschel} did not observe the entire sky. To maximize the number of clouds observed, the strategy followed by the GTO teams was to map the densest regions in the molecular clouds. These observations provide a unique high-resolution and high dynamic range view of the densest star-forming structures, in particular, for clouds far from the Galactic plane where the resolution of the NIR dust-extinction maps is limited. The obvious drawback of this choice is that the maps are incomplete, missing the extended low-column-density regions containing most of a cloud's mass, as seen in NIR extinction maps \citep[e.g.][]{2006A&A...454..781L, 2010A&A...512A..67L, 2008A&A...489..143L, 2011A&A...535A..16L} (Alves 2013, in prep.). In this paper we present a high-resolution, high dynamic range column-density map of the Orion complex using a combination of \textit{Planck} dust-emission maps, \textit{Herschel} dust-emission maps, and our own 2MASS NIR dust-extinction maps. The Orion column-density maps presented in this paper combine the robustness of the 2MASS NIR extinction maps with the resolution and coverage of the \textit{Herschel} and \textit{Planck} dust-emission maps and constitute the highest dynamic range column-density maps ever constructed for the entire Orion complex, covering $\SI{0.05}{mag} < A_K < \SI{10}{mag}$, or $\SI{1D21}{cm^{-2}} < N < \SI{2D23}{cm^{-2}}$. The Orion star-forming region, being the most massive and most active star-forming complex in the local neighborhood \citep[e.g.][]{1986ApJ...303..375M, 1991psfe.conf..125B, 1995A&A...300..903B, 2005A&A...430..523W, 2008hsf1.book..459B, 2011A&A...535A..16L}, is probably the most often studied molecular-cloud complex \citep[see][]{2008hsf1.book..459B, 2008hsf1.book..483M,2013ApJS..207...10R, 2013ApJ...766L..17S}. It contains the nearest massive star-forming cluster to Earth, the Trapezium cluster \citep[e.g.][]{1997AJ....113.1733H, 2000AJ....120.3162L, 2002ApJ...573..366M, 2012ApJ...748...14D}, at a distance of \SI{414}{pc} \citep{2007A&A...474..515M}). This paper is organized as follows. In Sect.~2 we briefly describe the data reduction process. Section~3 presents our approach to the problem of converting dust-emission into column-density. Section~4 is devoted to the application of the technique to the Orion~A and B molecular clouds. We discuss the results obtained in Sect.~5. Finally, in Sect.~6 we present a summary. We make use of PDF JavaScript to create figures with multiple layers: this make it easier to perform direct comparisons between different data or different results. Figures with multiple layers have buttons highlighted with a dashed blue contour in their captions. The hidden layers can be displayed only using a PDF reader with JavaScript enabled, such as Adobe$\textsuperscript{\textregistered}$ Acrobat$\textsuperscript{\textregistered}$, Foxit$\textsuperscript{\textregistered}$ Reader, or Evince. We also provide the hidden layers as separate figures in the appendix (in the electronic form of the journal). | \label{sec:conclusions} Our main results can be summarized in the following items: \begin{itemize} \item We presented optical-depth and temperature maps of the entire Orion molecular cloud complex obtained from \textit{Herschel} and \textit{Planck} space observatories. \item The maps have a \SI{36}{arcsec} resolution for \textit{Herschel} observations and a \SI{5}{arcmin} resolution elsewhere. In addition, we also produced a \SI{18}{arcsec} resolution optical-depth maps based on the SPIRE~250 data alone. \item We calibrated the optical-depth maps using 2MASS/\textsc{Nicest} extinction data, thus obtaining column-density extinction maps at the resolution of \textit{Herschel} with a dynamic range \SI{1D-2}{mag} to \SI{30}{mag} of $A_K$, or from \SI{4D20}{cm^{-2}} to \SI{6D23}{cm^{-2}}. \item We measured $C_{2.2}/\kappa_{850}$, that is, the ratio of the \SI{2.2}{\um} extinction coefficient and of the \SI{850}{\um} opacity. We found that the values obtained for both Orion~A and B cannot be explained using the \citet{1994A&A...291..943O} or the \citet{2001ApJ...548..296W} theoretical models of dust, but agree very well with the newer \citet{2011A&A...532A..43O} models for ice-covered silicate-graphite conglomerate grains. \item We examined the cumulative and differential area functions of the data, showing that over a large regime of extinction we observe a power-law $S(> A_K) \propto A_K^{-2}$, which is reminiscent of a simple isothermal model of molecular clouds; surprisingly, we do not see clear evidence of log-normality in the column-density pdf. \item We used the \textit{Planck/Herschel} maps to re-evaluate the \textit{local} Schmidt-law for star formation, $\Sigma_{YSO} \propto A_K^\beta$. We found that $\beta \simeq 2$ in Orion~A, confirming our earlier studies \citep{2013A&A...559A..90L, 2013ApJ...778..133L}. For Orion~B, we also found $\beta \simeq 2$, which is lower than our previous estimates as a result of the much improved angular resolution of the \textit{Herschel} observations. \end{itemize} | 14 | 4 | 1404.0032 |
1404 | 1404.2806_arXiv.txt | The Heterodyne Instrument for the Far Infrared (HIFI) on board the Herschel Space Observatory is composed of a set of fourteen double sideband mixers. We discuss the general problem of the sideband ratio (SBR) determination and the impact of an imbalanced sideband ratio on the line calibration in double sideband heterodyne receivers. The HIFI SBR is determined from a combination of data taken during pre-launch gas cell tests and in-flight. The results and some of the calibration artefacts discovered in the gas cell test data are presented here along with some examples of how these effects appear in science data taken in orbit. | \label{intro} The Herschel Space Observatory (\cite{pilbratt2010}) was launched on May 14th 2009 and successfully observed objects in the Sub-millimetre and Far Infrared (FIR) bands from the Solar System to the most distant reaches of the Universe. The mission ended on April 29$^{\rm th}$ 2013 when the Helium coolant boiled off. Its spectral coverage was provided by three instruments, PACS, SPIRE and HIFI. PACS (Photodetecting Array Camera and Spectrometer) was an imaging camera and low-resolution integral field spectrometer covering wavelengths from 55 to 210\,\si{\micro\metre} \cite{poglitsch2010}. SPIRE (Spectral and Photometric Imaging Receiver) was also a camera and a Fourier transform spectrometer (FTS) covering wavelengths from 194 to 671\,\si{\micro\metre} \cite{griffin2010}. HIFI (Heterodyne Instrument for the Far-Infra-red) was a heterodyne detector and spectrometer providing high resolution spectroscopy capability over two continuous frequency ranges of 488--1272 and 1430--1902\,GHz \cite{deGraauw2010}. This work addresses the calibration of the HIFI instrument. \par HIFI covers its spectral range using 14 heterodyne detectors, mixing down the FIR signal to radio frequencies. They are organised in 7 bands with 2 mixers each, sensitive to orthogonal polarizations. The mixers in each band are pumped by a pair of Local Oscillator (LO) chains, covering respectively the lower (LO chain {\it a}) and upper (LO chain {\it b}) frequencies of a band tuning range. Table \ref{mixer_overview} provides an overview of the frequency coverage and mixer technologies used in HIFI. \par \begin{table}[!t] \centering \begin{tabular}{|l|l|p{1cm}|l|l|l|l|} \hline Mixer & LO Frequency & IF BW & Detector & Beam & Feed and coupling\\ band & range (GHz)& (GHz) & technology & combiner & structure\\\hline \textbf{1} & 488--628 & 4--8 & SIS$^a$ \cite{delorme2005} & Beamsplitter& corrugated horn\\ & & & & microstrip & and waveguide\\\hline \textbf{2} & 634--794 & 4--8 & SIS$^a$ \cite{teipen2005} & Beamsplitter&corrugated horn\\ & & & & microstrip &and waveguide\\\hline \textbf{3} & 807--953 & 4--8 & SIS$^a$ \cite{delange2003} & Diplexer&corrugated horn\\ & & & & microstrip &and waveguide\\\hline \textbf{4} & 957--1114 & 4--8 & SIS$^a$ \cite{delange2003} & Diplexer&corrugated horn\\ & & & & microstrip &and waveguide\\\hline \textbf{5} & 1116--1272 & 4--8 & SIS$^b$ \cite{karpov2007} & Beamsplitter&lens and twin slot\\ & & & & microstrip & planar antenna\\\hline \textbf{6} & 1430--1698 & 2.4--4.8 & HEB$^c$ \cite{cherednichenko2008} & Diplexer&lens and twin slot\\ & & & & & planar antenna\\\hline \textbf{7} & 1701--1902 & 2.4--4.8 & HEB$^c$ \cite{cherednichenko2008} & Diplexer &lens and twin slot\\ & & & & & planar antenna\\\hline \end{tabular} \caption{Overview of mixer technology, materials and implemented antenna technology \cite{deGraauw2010}. $^a$ Nb-Al$_{2}$O$_{3}$-Nb, $^b$ Nb-AlN-NbTiN, $^c$ NbN Phonon cooled.} \label{mixer_overview} \end{table} \par Observing in the environment of space allows unobstructed by atmosphere coverage of the entire HIFI frequency range. It also removes the day night cycle constraints and allows for continuous visibility of a source over an observational day. From a calibration perspective, observing from space also gives a number of advantages over ground based telescopes. The absence of the atmosphere from the calibration equation removes one of the major sources of error in flux determination \cite{guan2012}. Owing to a careful thermal control of the detection chain, a high temperature stability is achieved, that reduces mixer gain variation, and returns better data quality with less baseline features such as standing waves and baseline distortions. However, at the shorter wavelengths of the Hot Electron Bolometer (HEB) mixers, data quality was at times poor due to system instability \cite{higgins2009, kooi2009}. \par In anticipation of this unique environment ambitious calibration accuracies were sought, with a baseline calibration uncertainty of 10\% and a goal calibration uncertainty of 3\% \cite{roelfsema2012}. The main sources of calibration error inherent to a double sideband heterodyne system were determined to be the sideband ratio (hereafter SBR), standing waves, and the calibration load coupling. Specific tests were implemented prior to launch to constrain these error sources. They are described in \cite{teyssier2008}. \par In this paper we discuss the sideband ratio derived from flight data and from the pre-launch test campaign conducted with a gas cell. Section \ref{side_band_ratio} gives the background on the significance of the sideband ratio and Section \ref{sbr_imbalance} describes the various phenomena involved in defining its characteristics. Section \ref{gascell_sideband_ratio} summarizes the results from the gas cell test campaign. Section \ref{side_band_ratio_in_data} provides examples of how the SBR manifests in certain areas of the HIFI frequency range. A discussion section follows summarizing the lessons learned from this calibration effort. | The gas cell test campaign provided thousands of measurement points prior to the launch of Herschel, building probably one of the largest dataset dedicated to the calibration of the sideband ratio in a double sideband receiver, and demonstrating the unique capabilities of HIFI as a high-resolution spectrometer already years before the Herschel launch. Yet, more than half a year after the mission completion the sideband ratio correction is still puzzling in several areas of the HIFI frequency range. One of the main reasons for this is the relative scarcity of measurement points that was inherent to the limited number of gases available for the gas cell tests. This frequency coverage granularity later proved to be insufficient in bands such as band 1 or 2 where steep sideband ratio variations have been confirmed to occur over relatively narrow frequency domains. \par The best way to join the dots between the pre-launch measurement points is to make use of high signal-to-noise spectral sweeps that can provide relative variation of line intensities within frequency steps of a few hundred MHz. The combination of the absolute sideband ratio points derived from the gas cell tests, and their relative evolution in flight data is therefore needed to derive the detailed structure of the gain profile over the whole HIFI range. This approach was considered before launch already when a full spectral scan of the methanol molecule was collected in the laboratory. While this dataset probably contains most of the answers we are after, it is to date still too complex to interpret due to often non-mature spectroscopic parameters (e.g. pressure broadening) for methanol, which is a mandatory input for a correct modelling of the molecular absorption (most of it is non-saturated) as a function of the frequency. \par We note that part of the complexity uncovered in e.g. band 1 was not anticipated, even during the design of the mixer circuits. From the early FTS measurements of the mixer broadband coupling, it was unclear that significant variations of the sideband ratio over frequency ranges as narrow as 4 GHz, such as the one evidenced e.g. around the $^{12}$CO 5-4 line (Fig. \ref{12co_54_plot}), would apply to the mixer response. In hindsight OCS measurements on a finer LO frequency grid would have helped unravel the coupling profile of bands 1 and 2. It is still unclear whether the discrepancy between those FTS measurements and the gas cell test data is simply due to the fact that the FTS tests did not use pumped mixers. We suggest that this discrepancy could be further investigated by groups involved in heterodyne detector design. \par One of the important questions for the astronomers is, what is the final absolute accuracy of the line intensity observed, which for HIFI often boils down to the contribution of the sideband ratio. Because the sideband ratio is currently assumed to be constant over most of the tuning range, the same is true for its accuracy, which is simply taken as the standard deviation of the sideband gains derived from the gas cell tests over a given band. We discussed in Section~\ref{sect_heb} that this is too pessimistic e.g. for the HEB bands. On top of that, our description of the sideband ratio is LO- and IF-dependent, hence the calibration accuracy should be treated accordingly. Other aspects of the sideband ratio calibration, such as standing wave and/or diplexer cross-talk (not discussed in detail in this paper), are too complex to measure at such granularity. This is where the need of a detailed instrument model becomes important, and such work is currently on-going within the HIFI team \cite{delforge2013}. It is unrealistic to assume that a theoretical model could accurately predict the frequency-dependent sideband ratio profile at any given LO frequency, so the idea here is not to correct observational data by synthetic instrument response function. Rather, such a model should provide the order of magnitude of typical calibration errors resulting from optical effects such as standing waves, esp. in the diplexer bands.\par An important lesson from the sideband ratio calibration effort is that a large fraction of the answers can be directly extracted from the science data themselves. While the gas cell test campaign offered measurement conditions with a fully controlled sky signal input, it was relatively constrained in time and performed at a moment when the instrument understanding was not fully mature, so some of the data peculiarities where still uncovered or not accounted for at the time. The regular science observations gathered over the almost four years of mission provide the complementary information that not only allows to revisit some of the laboratory data, but also to probe the instrument behaviour in an often larger variety of parameter space than could be considered pre-launch. The combination of the two is the key to build the most accurate picture of the receiver characteristics, and offer the best possible calibration to the legacy archive data.\par In that respect, the main source of information lies in the so-called spectral surveys, where a given spectral range is observed several times at nearby LO frequency tunings. This redundancy allows to invert the Double Sideband problem and build a Single Sideband solution - this is called {\it deconvolution} \cite{comito2002}. In this process, the gains applying respectively to the LSB and USB can be fitted in order to reconcile all individual DSB spectra with the SSB solution. Intrinsically the sideband gain profile inferred from this algorithm is only relative and needs to be tied to absolute values measured somewhere else (typically during the gas cell tests). The quality of the recovered sideband gain profile depends on the line density (if it is too large the risk of line blend is high, if it is too low the problem becomes degenerated at some frequencies). The HIFI team is currently running simulations on synthetic data with a variety of user-fed sideband ratio models and line density to identify the best data-set of spectral surveys to be used for that purpose, together with the accuracy associated with such an approach. The goal will be to extract an as continuous as possible profile of the sideband ratio over the HIFI observational range. Until then, the recommendation is to use the sideband ratio measured at particular line frequencies during the gas cell test campaign. This information, along with the calibration uncertainty associated with the sideband ratio, is communicated via the technical note\cite{teyssier2013} found on the HIFI calibration webpages\footnote{http://herschel.esac.esa.int/twiki/pub/Public/HifiCalibrationWeb/}. | 14 | 4 | 1404.2806 |
1404 | 1404.5617_arXiv.txt | We present here the first observationally based determination of the rate of occurrence of circumbinary planets. This is derived from the publicly available Kepler data, using an automated search algorithm and debiasing process to produce occurrence rates implied by the seven systems already known. These rates depend critically on the planetary inclination distribution: if circumbinary planets are preferentially coplanar with their host binaries, as has been suggested, then the rate of occurrence of planets with $R_p>6R_\oplus$ orbiting with $P_p<300$\ d is $10.0 ^{+18}_{-6.5}$\% (95\% confidence limits), higher than but consistent with single star rates. If on the other hand the underlying planetary inclination distribution is isotropic, then this occurrence rate rises dramatically, to give a lower limit of 47\%. This implies that formation and subsequent dynamical evolution in circumbinary disks must either lead to largely coplanar planets, or proceed with significantly greater ease than in circumstellar disks. As a result of this investigation we also show that giant planets (${>}10R_\oplus$) are significantly less common in circumbinary orbits than their smaller siblings, and confirm that the proposed shortfall of circumbinary planets orbiting the shorter period binaries in the Kepler sample is a real effect. | \label{sectIntro} In recent years the incredibly precise Kepler data has produced a wide range of important discoveries. Among these is the array of planets now known orbiting binary stars, proving not only that circumbinary planets can exist stably in such locations, but that they are not rare. At the time of writing seven systems with transiting circumbinary planets are known, being Kepler-16b \citep{Doyle:2011ev}, Kepler-34b and -35b \citep{Welsh:2012kl}, Kepler-38b \citep{Orosz:2012ip}, Kepler-47b and c \citep{Orosz:2012ku}, Kepler-64b/PH1 \citep{Kostov:2012vb,Schwamb:2012ts} and the recently published Kepler-413b \citep{Kostov:2014bx}. Although there are significant obstacles to routine detection in the form of large transit timing and duration variations \citep{Armstrong:2013kg}, the relatively small sample of Kepler eclipsing binaries has produced a sizeable number. Several questions remain: how abundant are these planets? How does the central binary affect their formation \citep[e.g.][]{Pelupessy:2013gl}? What evolutionary processes dominate in such an environment \citep[e.g.][]{Pierens:2008fq,Pierens:2013ee}? Some theoretical work has been done, showing that planet formation in a circumbinary disk could be hindered by raised planetessimal velocities \citep{Meschiari:2012er} over an area several AU in size, including the present orbits of the known planets. This implies that circumbinary planet formation may well proceed on wider orbits, with planets later migrating to their current positions \citep{Kley:2014vv}. The exact extent of the formation suppressing area is as yet unknown, and it has been proposed that planet formation in circumbinary disks may be helped by zones of lower velocity \citep{GMartin:2013be,Rafikov:2012wl}. How easily such planets form, and the evolutionary route they follow, represents an excellent constraint on planet formation in general. The Kepler sample of circumbinaries (CBs) has grown to a point where it can begin to tell us about these planets in general. Here we use it to extract what information we can on the rate of occurrence of CBs, as well as their distribution of inclinations. These are important indicators of the history of CB systems, showing whether formation proceeds easily of with difficulty, and whether scattering plays a key role in any subsequent evolution. We focus here on detached binary systems, and on planets with periods within 300\ d. For reasons of completeness we only utilise planets showing consecutive transits, i.e. those which produce a transit on each orbit. There are expected to be many `sparsely' transiting CBs that only occasionally transit \citep{Martin:2014ug}, and which can be expected to provide more information in future. We use occurrence rate here to mean the number of binaries with one (or more) planets as a fraction of the total binary number, leaving the question of multiple planets per binary to future work. | We have investigated the rates of occurrence of circumbinary planets orbiting close ($P_\textrm{bin}<$\mytilde60\ d) non-contact binary stars using the Kepler sample of eclipsing binaries. This produced a number of interesting results: \begin{enumerate} \item The most significant controlling distribution is that of planetary inclination - whether these planets lie preferentially coplanar with their host binaries, or in a more uniform pattern. Our results show that if such a uniform or even generally misaligned distribution is the norm, then the rate of occurrence of CBs must be exceptionally high, significantly more so than analogous rates for single stars. While not formally excluding very uniform, misaligned planetary inclination distributions, these results show that to exist such distributions need planetary formation rates at levels very difficult to explain. \item Conversely, if coplanarity is preferred, to the level implied by a Gaussian distribution with standard deviation \mytilde $5^\circ$ or tighter (although we note that the distribution by no means must be gaussian, and may even be bimodal) then the rate of occurrence of CBs is consistent with that of single star planets. Evidence suggests that circumbinary planets orbiting sub-AU binaries should be preferentially coplanar due to alignment of the protoplanetary disk, supporting this option \citep{Foucart:2013gk,Kennedy:2012jl}. \item CB giant planets (defined as $>=10R_\oplus$) are significantly less common than their smaller equivalents. There remains the possibility of a non-coplanar giant CB population at any rate of occurrence, formed for example by dynamical evolution, but a coplanar CB giant population on the same order as planets with $R<10R_\oplus$ is excluded, at least within our tested period range. Given that proto-planetary disk masses scale with the mass of the central object \citep{Andrews:2013ku}, and that more massive disks produce more gas giants \citep{Mordasini:2012ek} this supports the finding of \citet{Pierens:2008fq}, that CB Jupiter mass planets if present will likely orbit at larger distances from the central binary due to increased scattering. \item CB planets are less common in coplanar orbits around shorter period binaries ($P_\textrm{bin}<\mytilde\> 5-10$\ d) than around binaries of longer period. We have shown that this trend is not the result of sampling bias, with 99.9\% confidence for all tested misaligned planetary inclination distributions and 97.7\% for a coplanar distribution. The observed difference could be explained through a significantly different orbital distribution between planets orbiting shorter and longer period binaries (such as a more misaligned shorter population, so that we do not observe them) or by an effect of the formation of these binary systems \citep[see e.g.][]{Fabrycky:2007jh}. If shorter period binaries form through secular interactions with a tertiary stellar companion, planets in these systems would either be disrupted, or if present difficult to see due to dilution by the companion. If such close binaries have evolved to their current orbit via angular moment loss (through e.g. magnetic braking) then this process may influence the protoplanetary disk and impact planet formation. This remains a promising area of future work. \end{enumerate} To improve our knowledge of these unusual systems a larger sample of circumbinary planets needs to be found. Fortunately there are several possible routes to these discoveries, from searches for misaligned transiting planets to the use of radial velocities or binary eclipse timing. All of these will help to increase the sample size available, leading to new insights into their formation, evolution, and how these impact on general planet formation theories. The discovery of more misaligned planets will allow tighter constraints to be placed on planetary inclination distributions, answering questions about the dynamical evolution of these systems. Beyond this, future space missions such as PLATO and TESS should provide a great deal more new transiting, bright CB planets for further work. | 14 | 4 | 1404.5617 |
1404 | 1404.7228_arXiv.txt | In a newly born (high-temperature and Keplerian rotating) neutron star, $r$-mode instability can lead to stellar differential rotation, which winds the seed poloidal magnetic field ($\sim 10^{11}$ G) to generate an ultra-high ($\sim 10^{17}$ G) toroidal field component. Subsequently, by succumbing to the Tayler instability, the toroidal field could be partially transformed into a new poloidal field. Through such dynamo processes, the newly born neutron star with sufficiently rapid rotation could become a magnetar on a timescale of $\sim 10^{2-3}$ s, with a surface dipolar magnetic field of $\sim10^{15}$ G. Accompanying the field amplification, the star could spin down to a period of $\sim5$ ms through gravitational wave radiation due to the $r$-mode instability and, in particular, the non-axisymmetric stellar deformation caused by the toroidal field. This scenario provides a possible explanation for why the remnant neutron stars formed in gamma-ray bursts and superluminous supernovae could be millisecond magnetars. | Since the operation of the \textit{Swift} satellite, it has been widely suggested that the remnant compact objects formed in some gamma-ray bursts (GRBs) could be rapidly rotating, highly magnetized neutron stars (NSs). Such a millisecond magnetar scenario is helpful for understanding observations such as X-ray shallow decay and, in particular, plateau afterglows on timescales of $\sim 10^{2-4}$ s (e.g., Dai \& Lu 1998; Zhang \& M\'esz\'aros 2001; Fan \& Xu 2006; Yu et al. 2010; Metzger et al. 2011; Zhang 2013; Rowlinson et al. 2013) and the temporally ``extended'' gamma-ray emission on timescales of minutes of short GRBs (Gao \& Fan 2006; Metzger et al. 2008; Bucciantini et al. 2012). Recently, a similar energy source scenario was employed to interpret the high luminosity of some superluminous supernovae (e.g., Kasen \& Bildsten 2010) and was also suggested to power bright mergernova emission during the merger of a double NS system (Yu et al. 2013). In all of these cases, the high magnetic field of the NS is required to ensure that most of the rotational energy of the star can be released into the stellar wind in a sufficiently short time. Magnetars can generally be defined as special types of NSs with surface (though sometimes only interior) magnetic fields that are as high as $\sim 10^{14}$ G at least. The dissipation of the magnetic fields could power some high-energy electromagnetic emission, e.g., the GRB X-ray flares (Dai et al. 2006), and the bursts of soft gamma-ray repeaters (SGRs) and anomalous X-ray pulsars (AXPs; Thompson \& Duncan 1993). The strength of the surface dipolar magnetic fields of most Galactic SGRs and AXPs is indeed inferred to be on the order of $10^{14-15}$ G by $B_{p}=3.2\times10^{19}\sqrt{P\dot{P}}$, where $P$ is the pulse period and $\dot{P}$ is the period derivative (Olausen \& Kaspi 2013). However, the recently observed SGR 0418+5729 and Swift J1822.3--1606 both indicate much lower dipolar fields ($6\times 10^{12}$ G and $3.8\times 10^{13}$ G, respectively; Rea et al. 2013), which are typical for normal pulsars rather than magnetars. This ``contradiction'' hints that the high magnetic fields of some magnetars could be dominated by multipolar (e.g., toroidal) components, which are probably hidden in the stellar interior. The interior fields could be much stronger than those on the surface. The most straightforward consideration one might suggest is that the high magnetic field of a newly born magnetar may originate from the fossil magnetic fluxes in the progenitor core via the magnetic flux conservation, where the progenitor should be highly magnetized (Ferrario \& Wickramasinghe 2006). However, considering the possible existence of interior multipolar magnetic fields in SGRs/AXPs and the extremely rapid rotation of the GRB magnetars, it is believed that the origin of the high magnetic field is more likely to be associated with a dynamo process located deep in the stellar interior. Duncan \& Thompson (1992) proposed that an $\alpha-\omegaup$ dynamo could be supported by neutrino-driven turbulent convection and initially existing differential rotation in newly born millisecond NSs. Alternatively, in this Letter we suggest that the extremely rapid rotation of newly born NSs can spontaneously initiate a dynamo process via $r$-mode instability and the magnetic Tayler instability. In a rotating NS, $r$-modes arise due to the action of the Coriolis force with positive feedback (Andersson 1998; Friedman \& Morsink 1998). The growth of the $r$-modes can be suppressed by viscous damping and, in particular, by some non-linear effects. Specifically, by expanding the $r$-modes up to the second order of amplitude, the differential rotation induced by the modes can be found to determine a saturation state of the instability (S\'a \& Tom\'e 2005, 2006; Yu et al. 2009). As a result, a toroidal magnetic field component can be formed and amplified by winding up the seed poloidal field (Rezzolla et al. 2000; Rezzolla et al. 2001a, 2001b). As it increases in a stably stratified stellar interior, the toroidal field could enter into the Tayler instability and therefore can be partly transformed into a new poloidal component. Finally, a stable poloidal--toroidal twisted torus configuration appears in the stellar interior, which is connected by an enhanced dipolar field on the stellar surface (Braithwaite \& Spruit 2004). Such a dynamo mechanism has been previously investigated in the framework of accreting NS binaries (Cuofano \& Drago 2010; Cuofano et al. 2012), where the solid crust of the NSs can provide an extra effective suppression on the $r$-mode instability. In contrast, for a newly born NS, the crust cannot form initially due to the high stellar temperature. Moreover, the rotation of the newly born NS could be very close to the Keplerian limit, which is much more rapid than what an accreting binary NS can reach (Hessels et al. 2006). The chief purpose of this Letter is to report on the formation of the remnant magnetars harbored in GRBs. In the Section 2, we describe the model for the evolutions of $r$-modes, magnetic fields, and stellar rotations. Calculated results are presented in Section 3. Conclusion and discussions are given in Section 4. | \label{Sec V} By considering the differential rotation caused by $r$-mode instability in a newly born, rapidly rotating NS, we calculate the evolution of the stellar magnetic fields, where an ultra-strong toroidal magnetic field is generated. Succumbing to the Tayler instability, the toroidal field is partially transformed into a new poloidal field. Through such dynamo processes, the NS could become a magnetar with a surface dipolar field of a strength $\sim 10^{15}$ G on timescales $\sim 10^{2-3}$ s, the precondition of which is that the NS should rotate initially with a nearly Keplerian period, $P_{\rm i}\lesssim 1.7$ ms. Such a condition could easily be satisfied in the situation of GRBs. For somewhat longer periods, $1.7{\rm ms}<P_{\rm i}\lesssim3$ ms, this dynamo could work in principle, but the strengths of the fields become much lower. Moreover, the amplification of the surface field is delayed to a very late time, at which more complexity (e.g., the formation of a crust) is involved. In any case, the long time delay could make the model inapplicable for GRB magnetars. Finally, for $P_{\rm i}>3$ ms, the dynamo processes would never happen and a normal magnetic field keeps in the NS, because the $r$-mode instability is suppressed by viscosities. Due to the magnetic dissipation, the $r$-mode-induced GW radiation becomes very weak. Alternatively, another strong GW radiation is produced due to the high deformation of the NS by the toroidal magnetic field, which could cause the star to be a promising target for GW detection. As a result, accompanying the magnetic field amplification, the spin periods of GRB magnetars would be increased to $\sim 5$ ms. In other words, the ``initial" spin periods derived from GRB afterglow observations should be basically consistent with such a value. Furthermore, due to the GW radiation, a remarkable amount of the rotational energy of the NS can be released into the GW. Therefore, the supernova remnant around the magnetar cannot be as highly energized as usually considered. In observation, analysis of the X-ray spectra of some supernova remnants associated with magnetar candidates Vink \& Kuiper (2006) revealed that the total energy in these supernova remnants is almost nothing, which may favor our model. In other words, some Galactic magnetars may share the same origin mechanism presented here. | 14 | 4 | 1404.7228 |
1404 | 1404.6704_arXiv.txt | Evidence from the BICEP2 experiment for a significant gravitational-wave background has focussed attention on inflaton potentials $V(\phi) \propto \phi^\alpha$ with $\alpha=2$ (``chaotic'' or ``$m^2\phi^2$'' inflation) or with smaller values of $\alpha$, as may arise in axion-monodromy models. Here we show that reheating considerations may provide additional constraints to these models. The reheating phase preceding the radiation era is modeled by an effective equation-of-state parameter $w_{\rm re}$. The canonical reheating scenario is then described by $w_{\rm re}=0$. The simplest $\alpha=2$ models are consistent with $w_{\rm re} = 0$ for values of $n_s$ well within the current $1\sigma$ range. Models with $\alpha=1$ or $\alpha=2/3$ require a more exotic reheating phase, with $-1/3<w_{\rm re}<0$, unless $n_s$ falls above the current $1\sigma$ range. Likewise, models with $\alpha=4$ require a physically implausible $w_{\rm re}>1/3$, unless $n_s$ is close to the lower limit of the $2\sigma$ range. For $m^2\phi^2$ inflation and canonical reheating as a benchmark, we derive a relation $\log_{10}\left(T_{\rm re}/10^6\,{\rm GeV} \right) \simeq 2000\,(n_s-0.96)$ between the reheat temperature $T_{\rm re}$ and the scalar spectral index $n_s$. Thus, if $n_s$ is close to its central value, then $T_{\rm re}\lesssim 10^6$~GeV, just above the electroweak scale. If the reheat temperature is higher, as many theorists may prefer, then the scalar spectral index should be closer to $n_s\simeq0.965$ (at the pivot scale $k=0.05\,{\rm Mpc}^{-1}$), near the upper limit of the $1\sigma$ error range. Improved precision in the measurement of $n_s$ should allow $m^2\phi^2$, axion-monodromy, and $\phi^4$ models to be distinguished, even without precise measurement of $r$, and to test the $m^2\phi^2$ expectation of $n_s\simeq0.965$. | 14 | 4 | 1404.6704 |
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1404 | 1404.1915_arXiv.txt | We develop a theoretical framework that extracts a deeper understanding of galaxy formation from empirically-derived relations among galaxy properties by extending the main-sequence integration method for computing galaxy star formation histories. We properly account for scatter in the stellar mass--star formation rate relation and the evolving fraction of passive systems and find that the latter effect is almost solely responsible for the age distributions among $z\sim0$ galaxies with stellar masses above $\sim 10^{10}\,\msun$. However, while we qualitatively agree with the observed median stellar metallicity as a function of stellar mass, we attribute our inability to reproduce the distribution in detail largely to a combination of imperfect gas-phase metallicity and $\alpha$/Fe ratio calibrations. Our formalism will benefit from new observational constraints and, in turn, improve interpretations of future data by providing self-consistent star formation histories for population synthesis modeling. | The study of galaxy evolution is swimming in a flood of new multi-wavelength data. Recent observations have characterized (1) the evolution of stellar mass function \citep[e.g.,][]{Bielby12, Muzzin13}; (2) the bimodal nature of galaxies into quiescent and star-forming to $z\sim3$ \citep[e.g.,][]{Brammer11, Muzzin13}; (3) the evolving correlation between the stellar masses and star formation rates of star-forming systems \citep[e.g.,][]{Noeske07, Daddi07, Karim11, Whitaker12, Gonzalez14}; (4) a tight ``fundamental metallicity relation" (FMR) among stellar mass, star formation rate, and gas-phase metallicity \citep[e.g.,][]{KE08, LaraLopez10, Mannucci10, AM13}; (6) the structure of cold gas fueling galaxies \citep[e.g.,][]{Genel10}; and (6) the environmental influence on galactic properties \citep[e.g.,][]{Peng10, Pasquali10, Lin14}. Yet, these various empirical relations beg a theoretical framework that answer questions about their effects on and relative importance in the buildup of stellar mass in the universe. For example, what role does the quiescent phase of galaxy evolution play in setting the growth of the stellar population, and do clumps in cold streams or environmental processes fundamentally drive star formation? Moreover, understanding the history of star formation in the universe is critical, not only for its own sake, but because the knowledge feeds back into the interpretation of the torrent of extragalactic observations. Extracting most galaxy properties from the data, for example, requires stellar population synthesis models for which the star formation history is a necessary input \citep[e.g.,][]{Tinsley80,Leitherer99, Schaerer13}. These histories are also central to understanding the nature of the mass-metallicity relation \citep[e.g.,][]{PS13} and the evolution of the law relating star formation rate to gas density \citep[e.g.,][]{Bigiel08}. To avoid the degeneracies and sub-grid prescriptions involved in formulating a completely theoretical model of galaxy formation and evolution for comparison with observations, we are motivated to organize the empirical data itself into a cohesive structure describing the buildup of stellar mass within galaxies. Such a path is provided by the Main Sequence Integration (MSI) method \citep[e.g.][]{Leitner12}. This scheme traces the star formation history of a mock galaxy by using the observed average star formation rate as a function of stellar mass and redshift to inform the growth of the system in each time step. It can also track the buildup of stellar metallicity by incorporating the empirically-determined FMR \citep{PS13}. However, the current method ignores the significant scatter in these empirical relations \citep[e.g.,][]{Noeske07, Daddi07, Karim11, Guo13, Gonzalez14} as well as the quenching of star formation implied by the evolving fraction of quiescent galaxies. Both effects contribute to large variations in the star formation history from system-to-system, which in turn, imprint on the derivations of galaxy properties from the data. To study the influence of empirical phenomenology on the star formation process more fully, we develop an improved MSI model that introduces fluctuations into the star formation process and, by construction, reproduces both the scatter in the stellar mass--star formation rate relation and the evolving fraction of quiescent galaxies. In \S\ref{sec:basic}, we review the basic MSI method, which traces the mean history of smoothly and continuously star-forming systems. We present our new prescriptions for generating variations in the star formation history, both in time and from galaxy-to-galaxy, in \S\ref{sec:scatt}, and test our method using Sloan Digital Sky Survey (SDSS) observations from \citep{Gallazzi05} of the scatter in galaxy ages and stellar metallicities in \S\ref{sec:results}. Finally, we discuss and summarize our conclusions in \S\ref{sec:forward}. Throughout, we adopt a $\Lambda$CDM cosmology with $\omm=0.28$, $\oml=0.72$, and $H_0=70\,{\rm km/s/Mpc}$ and a \citet{Chabrier03} initial mass function (IMF). Furthermore, $\log$ always refers to the base-10 logarithm. | \label{sec:forward} In this study, we develop a cohesive theoretical formalism for translating empirical relations into an understanding of the variations in galactic star formation histories. We achieve this goal by incorporating into the Main Sequence Integration method the scatter suggested by the evolving fraction of quiescent galaxies and the spread in the observed stellar mass--star formation rate relation. We find excellent agreement between our model results and the \citet{Gallazzi05} measurements of galactic ages---and approximate agreement with their median stellar metallicities---as a function of stellar mass but are unable to produce the observed variations among the stellar metallicities of dwarf galaxies (see Appendix for information about our tabulated results). Our work confirms that quenching is the key factor in determining the age distribution of massive systems. Because our implementation of quenching is stochastic and does not account for the frequency of different environments, environmental effects likely do not contribute very significantly; observational correlations between environment and age may simply result from the strong clustering of massive galaxies into overdense regions. On the other hand, among star-forming galaxies, variations in star formation rate likely correlate on time-scales longer than $10^{8}\,\yr$---unless additional starbursts strongly influence dwarf galaxy formation \citep[e.g.][but see also \citealt{GarrisonKimmel13}]{Weisz12, Amorisco14, Madau14}---and probably around $10^{9}\,\yr$, which suggests that environmental effects could play some role in setting star formation rates. However, additional starbursts beyond the normal scatter reflected in observed stellar mass--star formation rate relation are not required to explain the distribution of stellar age as a function of mass. Further, our work may reveal missing ingredients in the understanding of galaxy metallicities. For example, we find that results using representative formulations of the fundamental relation among gas-phase metallicity, stellar mass, and star formation rate approximately bracket the observed median stellar metallicities. Thus, we conclude that the uncertain calibration of gas-phase metallicity is at least partially responsible for our inability to reproduce the observations in detail. However, properly modeling the $\alpha$/Fe ratio as a function of the age of a stellar population also promises to improve significantly our comparisons with the data---particularly in generating the stellar metallicity scatter among dwarf galaxies, which retain less than $5\%$ of the metals they have ever made in their stars \citep{Peeples14}---and our understanding of the buildup of the stellar metallicity relation through time. Moreover, directly connecting empirical phenomenology with our framework for galaxy evolution gives us deeper insights into both data and theory. For example, we can improve the derivation of galaxy properties from future observations with our MSI-formalism, iteratively solving for the stellar masses and star formation rates, with self-consistent star formation histories. Such a procedure would yield more accurate results than would be achievable by assuming, e.g., a simple, exponentially-declining star formation rate in the required population synthesis calculation. Additionally, with regard to theoretical models, our framework points to degeneracies in the testing of numerical simulations and semi-analytic prescriptions against observations and suggests that physical processes correlating star formation on different time-scales may be distinguished by the scatter in the resulting galactic ages of dwarf galaxies. Thus, our framework clearly provides a useful organizational structure through which the behavior of galaxy properties can answer questions about the star formation and enrichment history of universe. These answers will increasingly clarify the field of galaxy formation as new streams of data continue to flow. | 14 | 4 | 1404.1915 |
1404 | 1404.6018_arXiv.txt | Spectral type recognition is an important and fundamental step of large sky survey projects in the data reduction for further scientific research, like parameter measurement and statistic work. It tends out to be a huge job to manually inspect the low quality spectra produced from massive spectroscopic survey, where the automatic pipeline may not provide confident type classification results. In order to improve the efficiency and effectiveness of spectral classification, we develop a semi-automated toolkit named ASERA, \emph{A} \emph{S}pectrum \emph{E}ye \emph{R}ecognition \emph{A}ssistant. The main purpose of ASERA is to help the user in quasar spectral recognition and redshift measurement. Furthermore it can also be used to recognize various types of spectra of stars, galaxies and AGNs (Active Galactic Nucleus). It is an interactive software allowing the user to visualize observed spectra, superimpose template spectra from the Sloan Digital Sky Survey (SDSS), and interactively access related spectral line information. It is an efficient and user-friendly toolkit for accurate classification of spectra observed by LAMOST (the Large Sky Area Multi-object Fiber Spectroscopic Telescope). The toolkit is available in two modes: a Java standalone application and a Java applet. ASERA has a few functions, such as wavelength and flux scale setting, zoom in and out, redshift estimation, spectral line identification, which helps user to improve the spectral classification accuracy especially for low quality spectra and reduce the labor of eyeball check. The function and performance of this tool is displayed through the recognition of several quasar spectra and a late type stellar spectrum from the LAMOST Pilot survey. Its future expansion capabilities are discussed. | \label{Introduction} The Large Sky Area Multi-object Fiber Spectroscopic Telescope (LAMOST) is a special reflecting Schmidt telescope specialized for conducting spectroscopic surveys with a wide field of view and a large aperture \citep{Wang:96}. One of the key scientific goals of LAMOST is the extragalactic spectroscopic survey of the large scale structure of the Universe and the physics of galaxies and quasars \citep{2011AAS...21812307W}. The pilot survey \citep{2012RAA....12.1197C} performed from October 2011 to June 2012 and the regular survey started in September 2012. There have been already millions of targets observed, including thousands of quasar candidates. Then to recognize quasars via spectra becomes essential for critical candidate confirmation and follow up scientific work. Spectra with high signal-to-noise ratio (SNR) are easily classified and the physical parameters can be determined with high accuracy using the LAMOST data processing pipeline \citep{2012RAA....12.1243L}. However there are still a large number of spectra with low SNR and probably some defects (e.g., skylight residual, splice connecting red part and blue part). Before the automatic pipeline is upgraded to be intelligent enough, eyeball check is in great need and enough astronomical knowledge is necessary. In SDSS quasar survey, visual inspection has been widely used to ensure the reliability of spectral identifications (P$\hat{a}$ris et~al. 2012). However in a large sky spectra survey, the quantity of spectra is very large. In order to reduce human efforts, we developed a spectrum eye recognition toolkit which provides a flexible platform to help identifying quasar spectra and estimating their redshifts at the same time. We call this toolkit ASERA. Both fits-formatted and image-formatted spectral files are supported. The input source can be placed in a local storage device, or distributed on the internet, described by a URL name. Since the software is developed using the Java programming language, it can be either started as a desktop application or accessed via a web browser, after deploying it as a Java applet. ASERA is initially dedicated to experienced spectrum analysts. It can also be used by teachers, undergraduate students and amateur astronomers. In the following sections, we first describe the detailed design and realization of the toolkit. Then several examples and figures are presented to explain how to use this toolkit on the spectra from LAMOST. In the end we discuss the system error on the estimated redshift $z$ and the following upgrade plans to extend the functionality of ASERA. | To improve the efficiency and effectiveness of spectral classification, ASERA, a spectrum eye recognition assistant, is developed using Java programming language, especially designed for quasar spectral recognition. The toolkit includes a graphical interactive interface with both the target spectrum and the template spectrum plotted, a group of convenient viewport adjustment functions to provide entire or partial inspection of the spectrum arbitrarily, and various spectral templates helping users to identify the target spectrum by eye. Via choosing a suitable redshift $z$ interactively, an artificial spectrum can be generated from a composite spectrum from Sloan Digital Sky Survey (SDSS). By comparing the generated spectrum with the target spectrum, taking the human experience as reference, users can finally recognize whether the target spectrum is a quasar or not, without being hampered by the partial abnormal or low SNR spectra. At the same time, ASERA may estimate the redshift $z$ of the recognized quasar spectrum. Several quasar spectra from the LAMOST Pilot survey are tested to show the advantage of this toolkit in handling low SNR spectra with skylight residual or stray light. ASEAR can be used to recognized various types of stars, galaxies and AGNs by importing their related template spectra. The systematic error of the redshift calculation is discussed. The toolkit will be publicly available as soon as possible and user may contact the author for a trial edition at present. In the future, FITS spectral files besides LAMOST and SDSS, will be supported further. Also, we will realize the access to the online spectral service using the Simple Spectral Access Protocol (SSAP) proposed by the IVOA. In addition, spectra in VOTable format will also be recognized and processed. | 14 | 4 | 1404.6018 |
1404 | 1404.7855_arXiv.txt | The large tensor spectrum recently observed by the BICEP2 Collaboration requires a super-Planckian field variation of the inflaton in the single-field inflationary scenario. The required slow-roll parameter $\epsilon\approx 0.01$ would restrict the e-folding number to around 7 in (sub-)Planckian inflationary models. To overcome such problems, we consider a two-field scenario based on the natural assisted supersymmetric (SUSY) hybrid model (``natural SUSY hybrid inflation'' \cite{NHinf}), which combines the SUSY hybrid and the natural inflation models. The axionic inflaton field from the natural inflation sector can admit the right values for the tensor spectrum as well as a spectral index of $0.96$ with a decay constant smaller than the Planck scale, $f\lesssim M_P$. On the other hand, the vacuum energy of $2\times 10^{16} \gev$ with 50 e-folds is provided by the inflaton coming from the SUSY hybrid sector, avoiding the eta problem. These are achieved by introducing both the U(1)$_R$ and a shift symmetry, and employing the minimal K$\ddot{\rm a}$hler potential. | Cosmological inflation not only resolves the problems in the standard big bang cosmology such as the homogeneity and flatness problems, but also explains the cosmological perturbations in matter density and spatial curvature \cite{review}. Those could naturally arise from the vacuum fluctuations of light scalar field(s) during inflation and be promoted to classical fluctuations around the time of the horizon exit. Indeed, the primordial power spectrum generated by an inflaton \cite{InflationFluctuation} turned out to be quite consistent with the cosmic microwave background (CMB) observations \cite{COBE,Komatsu:2010fb,Ade:2013zuv}. At the present, the inflationary paradigm seems to be strongly supported by various cosmological observations. Recently, the BICEP2 Collaboration reported their observation on the B mode of the primordial gravitational wave \cite{Ade:2014xna}. Their measurement indicates a large tensor spectrum or a large tensor-to-scalar ratio in the power spectrum \cite{Ade:2014xna}: \dis{ r=0.2^{+0.07}_{-0.05}\quad ({\rm or}~~ 0.16^{+0.06}_{-0.05} ~), } (after foreground subtraction with the best dust model). The tensor spectrum is evaluated on the scale, $x_{\rm ls}/100 \lesssim k^{-1} \lesssim x_{\rm ls}$ with $x_{\rm ls} = 14,000 \mpc$, which corresponds to a change in the e-folding number, $\Delta N \sim 4$. For smaller scales, however, the tensor spectrum by the primordial gravitational wave is suppressed. Moreover, it is not constrained by observation anymore. Such a large tensor-to-scalar ratio for the large scale implies relatively large values of the slow-roll parameter $\epsilon$ and the vacuum energy during that period of inflation: \dis{ \label{imply} \epsilon^*\approx 0.01 \quad {\rm and}\quad V^{1/4}\approx 2.08\times 10^{16} \gev . } In the slow-roll regime, the large $r$ requires a {\it super}-Planckian field variation for the inflaton in single-field inflationary models \cite{Lyth:1996im,lyth2}: \dis{ \label{lythBD} \frac{\Delta \varphi}{\Mp} \gtrsim \mathcal{O} (1) \times \bfrac{r}{0.1}^{1/2}, } where $M_P$ denotes the reduced Planck scale ($\approx 2.4\times 10^{18} \gev$). However, a super-Planckian field variation might imply the breakdown of an effective field theory description on inflation. On the other hand, the problem of {\it sub}-Planckian inflation is that such a relatively large $\epsilon \approx 0.01$ yields a too small e-folding number: \dis{ \label{efoldBD} \Delta N \approx \frac{1}{\Mp} \int \frac{d\varphi}{\sqrt{2\epsilon}} \approx 7 ~\bfrac{\Delta \varphi}{\Mp} \sqrt{\frac{0.01}{\epsilon}} . } Thus, only $\Delta N \sim 7$ is maximally obtained for $\Delta \varphi\sim M_P$. In order to get a large enough e-folding number, hence, either the field value must be super-Planckian or the slow-roll parameter $\epsilon$ should somehow be made to rapidly decrease (after about 7 e-folds \cite{Choi:2014aca}).\footnote{For the possibility of a large tensor-to-scalar ratio with single-field inflation models, see Refs.~\cite{Choudhury:2014kma,Antusch:2014cpa}.} For super-Planckian field values, however, it is necessary to invoke a symmetry to avoid the disastrous higher-order terms~\cite{Freese:1990rb,Kim:2014dba}. For a too rapidly decreasing $\epsilon$, the power spectrum may violate the observational constraint. To be consistent with the observation of the CMB, the power spectrum should be maintained as almost a constant within the observable scales in the CMB, which is within the range of $10 \mpc \lesssim k^{-1} \lesssim x_{\rm ls}$, corresponding to to $\Delta N\sim 7$ \cite{Ade:2013zuv}. For smaller scales, the power spectrum is constrained only to be smaller than around $10^{-2}$ by the argument of the missing primordial black hole \cite{Khlopov:1980mg}, or $10^{-4}$ by acoustic damping \cite{Chluba:2012we,Khatri:2013dha}, and especially less than 0.007 for $10^{-5} \mpc \lesssim k^{-1} \lesssim 10^{-4} \mpc $ from big bang nucleosynthesis (BBN)~\cite{Jeong:2014gna}. There is also a stronger constraint, ${\cal P}_\zeta \lesssim 10^{-6}$, for a specific type of dark matter using the nonobservation of the ultracompact mini halo \cite{Bringmann:2011ut}.\footnote{The improved observation on the spectral distortions on the CMB will constrain the power spectrum more strongly in the large scales corresponding to $\Delta N\sim 17$ from $x_{\rm ls}$ \cite{Chluba:2012we,Khatri:2013dha}.} Actually, it is quite hard to accommodate all the above stringent constraints within the single-field inflationary framework. The inflationary scenario is, indeed, based on the quantum theory of scalar fields. For successful inflation, the inflaton mass is required to be much lighter than the Hubble scale during inflation. As seen in the Higgs boson and the gauge hierarchy problem in elementary particle physics, however, it is highly nontrivial to keep a small enough inflaton mass against quantum corrections. Actually, only a few ways to get a light scalar are known in quantum field theory: i.e., by introducing {\bf (i)} the supersymmetry (SUSY), {\bf (ii)} a global U(1) symmetry, or {\bf (iii)} a strong dynamics. We will discuss here only the first two possibilities for a small inflaton mass. SUSY is an excellent symmetry that can protect a small scalar mass against quantum corrections. However, the problem in SUSY inflationary models is that SUSY must be broken due to positive vacuum energy of the Universe during inflation, even if it was introduced: positive vacuum energy inflating the Universe can induce a Hubble scale inflaton mass in supergravity (SUGRA), violating a slow-roll condition, $\eta\sim {\cal O}(1)$. It is called the ``eta ($\eta$) problem'' in SUGRA inflation models. It is a quite generic problem requiring an elaborate model construction to overcome. In SUSY hybrid inflation or ``$F$-term inflation'' \cite{FtermInf2,FtermInf}, fortunately the Hubble induced mass term is accidentally canceled out with the minimal K$\ddot{\rm a}$hler potential and the Polonyi-type superpotential during inflation. The specific form of the superpotential can be guaranteed by the introduced U(1)$_R$ symmetry. In this class of models, SUSY-breaking positive vacuum energy generates a logarithmic quantum correction to the constant scalar potential, which can draw the inflaton to the true minimum, triggering reheating of the Universe by the waterfall fields. The waterfall fields develop nonzero vacuum expectation values (VEVs) at their true minima, which can be determined with CMB anisotropy \cite{FtermInf}. They turn out to be tantalizingly close to the scale of the grand unified theory (GUT), ($\approx 2\times 10^{16} \gev$). Due to this, the waterfall fields can be regarded as GUT-breaking Higgs bosons in this class of models \cite{3221,422,FlippedSU(5),SO(10)}. In the (original) SUSY hybrid inflation model, a red-tilted power spectrum \cite{FtermInf} around \begin{equation} n_\zeta \approx 1+2\eta\approx 1-\frac{1}{N_e}\approx 0.98 \end{equation} is predicted for $N_e=50$ -- 60 e-folds. It is too large compared to the present bound on the spectral index. On the contrary, the tensor spectrum is too small to detect, $r\lesssim 0.03$ in this class of models \cite{Shafi:2010jr,Rehman:2010wm,Okada:2011en,Civiletti:2011qg}. Basically, the required slow-roll parameter, $\epsilon\approx 0.01$, is hard to get in SUSY hybrid models. In the ``natural inflation'' model \cite{Freese:1990rb}, the inflaton is looked upon as a pseudo-Nambu-Goldstone boson introduced from the spontaneous breaking mechanism of an anomalous global U(1) symmetry, U(1)$_{\rm PQ}$. By instanton effects, which break U(1)$_{\rm PQ}$ to a shift symmetry, a sinusoidal-type inflaton potential can be generated in this model. Because of a remaining shift symmetry, the inflaton field does not appear in the K$\ddot{\rm a}$hler potential in the SUSY version of the natural inflation model \cite{NaturalChaotic}. As a result, the unwanted Hubble scale inflaton mass term is not induced in the SUGRA potential during inflation. For a small enough slow-roll parameter $\eta$, however, the U(1)$_{\rm PQ}$ breaking scale or the ``axion decay constant'' $f$ must be larger than the Planck scale, \begin{equation} f \gtrsim 3 M_P ~. \end{equation} This implies that U(1)$_{\rm PQ}$ should be valid above the Planck scale. However, such a U(1)$_{\rm PQ}$ is not natural, because quantum gravity effects are known to break all continuous global symmetries, including U(1)$_{\rm PQ}$. One possible way to obtain an effectively large $f$ from a sub-Planckian Peccei-Quinn scale is to employ multiple axionic inflaton fields \cite{Kim:2004rp,doubleAxion}. Interestingly enough, the ``natural'' assisted SUSY hybrid inflation model or ``natural hybrid (NH) inflation'' \cite{NHinf}, which combines the SUSY hybrid and the natural inflation models, can cure the problems of the SUSY hybrid and natural inflationary models by interplay of the two inflatons: it can yield the desired value $n_\zeta\approx 0.96$ and realize $f\ll M_P$. The NH inflation introduces a shift and U(1)$_R$ symmetries, and it provides two light inflaton fields: one is the SUSY hybrid inflaton, and the other is the axionic one. Their Hubble induced masses are not generated if the K$\ddot{\rm a}$hler potential is of the minimal form \cite{NHinf}. Thus, the $\eta$ problem can be avoided. The smallness of the higher-order terms (particularly the quartic coefficient) in the K$\ddot{\rm a}$hler potential, however, needs to be justified with a UV quantum gravity theory. Since just a shift symmetry rather than U(1)$_{\rm PQ}$ is employed, the decay constant $f$ here does not have to be associated with a spontaneous symmetry-breaking mechanism. In this paper, we will consider a two-field inflationary scenario \cite{Choi:2012hea} by the NH inflation model to account for the observed large tensor spectrum: we will attempt to show that the NH inflation model achieves the large tensor-to-scalar ratio $0.16$ by interplay of the two inflaton fields. A large enough e-folding number during inflation to resolve the homogeneity and flatness problems can be obtained with large vacuum energy before the waterfall field destabilizes the vacuum. It is mainly driven by the inflaton coming from the SUSY hybrid model. On the other hand, the cosmological observables associated with quantum fluctuations of the inflaton are mainly provided by the inflaton of the natural inflation. During the first seven e-folds after the comoving scale of the last scattering exits the horizon, inflation is driven by the two inflaton fields, explaining the observed spectra in the CMB. After the axionic inflaton reaches almost the minimum of its potential, however, the inflationary history follows that of the SUSY hybrid models until the waterfall fields terminate the inflation, eventually yielding about 50 e-folds \cite{doublehybrid}. In this model, the field values of the SUSY hybrid inflaton during inflation are sub-Planckian ($\sim 10^{16-17} \gev$), while the axionic inflaton moves a Planckian distance along the phase direction \cite{phase}, which would not cause harm. This paper is organized as follows. In Sec. \ref{sec:model}, we set up the SUGRA model. In Sec. \ref{sec:Inf}, we discuss the spectrum and its indices for both scalar and tensor perturbations in the NH inflation model, which are the main results. We conclude in Sec. \ref{sec:conclusion}. | \label{sec:conclusion} The large tensor-to-scalar ratio, $r\approx 0.16$ observed by the BICEP2 Collaboration requires a super-Planckian field variation during the inflationary era in the single-field inflation scenario, which might destroy field theory description on inflation. Although it could be controlled using a shift symmetry, a super-Planckian decay constant would be another obstacle for constructing an inflation model based on effective field theory. In this paper, we have considered a two-field inflationary scenario based on the NH inflationary model, which combines the natural and SUSY hybrid inflation models. In this model, the Hubble scale induced mass term for the two inflaton fields, $\sigma$ and $a$ can be avoided by introducing the U(1)$_R$ and a shift symmetries, and employing the minimal form of the K${\rm \ddot{a}}$hler potential. The inflation is terminated by the VEVs of the waterfall fields, which is around the GUT scale. The power spectrum in the observable scale of $\Delta N \sim 8$ is determined mainly by the axionic inflaton field, while the needed e-folding number can be obtained from the hybrid inflation sector before the waterfall field terminates inflation. We find that the large tensor spectrum corresponding to $r\approx 0.16$ as well as the spectral index of the scalar power spectrum $n_\zeta\approx 0.96$ are well obtained with a sub-Planckian decay constant, $f\lesssim M_P$. While the field values of the SUSY hybrid inflaton $\sigma$ are sub-Planckian ($\sim 10^{16-17} \gev$) during inflation, the axionic inflaton $a$ moves a Planckian distance along the phase. However, the cosmological observables are determined while it stays in the sub-Planckian regime. | 14 | 4 | 1404.7855 |
1404 | 1404.5850_arXiv.txt | We present {\kep} and {\swift} observations of StHa 169 which is currently classified as a symbiotic binary. The {\kep} light curve shows quasi periodic behaviour with a mean period of 34 d and an amplitude of a few percent. Using {\swift} data we find a relatively strong UV source at the position of StHa 169 but no X-ray counterpart. Using a simple two component blackbody fit to model the combined {\swift} and 2MASS spectral energy distribution and an assessment of the previously published optical spectrum, we find that the source has a hot ($\sim$10,000K) component and a cooler ($\sim$3700K) component. The {\kep} light is dominated by the cool component and we attribute the variability to pulsations in a red giant star. If we remove this approximate month long modulation from the light curve, we find no evidence for additional variability in the light curve. The hotter source is assigned to a late B or early A main sequence star. We briefly discuss the implications of these findings and conclude that StHA 169 is a red giant plus main sequence binary. | Symbiotic stars are interacting binary systems containing a red giant star and a hotter component, which can be a white dwarf, a main sequence star or even a neutron star (see Mikolajewska 2007 for a review). A relatively small fraction of these binaries show evidence for accretion onto the hot component via a disc, while the remainder show evidence of accretion via the wind from the giant star and, in some systems, nuclear burning occurs on the surface of the hot component (see Kenyon \& Webbink 1984). Some systems such as CH Cyg, have produced jets (e.g. Taylor, Seaquist \& Mattei 1986, Crocker et al. 2001, Galloway \& Sokolski 2004) and large variations ($\sim$5 mag) in optical brightness over year long timescales (e.g. Mikolajewski, Mikolajewska \& Khudiakova 1990). More recently, evidence has been presented which suggests that symbiotic stars could be progenitors of a fraction of supernovae 1a explosions (e.g. Dilday et al. 2012). Given the fact that symbiotic stars contain a red giant star, the binary orbital periods are the longest (ranging from several hundreds of days to many years) found in the many different types of interacting binaries. As such, long-duration photometric surveys such as OGLE have been used to search for signatures of the binary period (e.g. Angeloni et al. 2014). In contrast, high time resolution photometry or spectroscopic observations are required to search for the presence of phenomena such as accretion whose rate can be variable over the medium to long timescale (e.g. Sokolski, Bildstein \& Ho 2001 and Sokolski \& Kenyon 2003). The {\kep} misson (Borucki et al. 2010) provides a unique opportunity to study objects such as symbiotic binaries on short timescales (1 min) and also much longer timescales (the initial {\kep} pointing lasted approximately 4 years). There were two objects classified as symbiotic binaries in the {\kep} field -- the previously mentioned system CH Cyg -- and StHA 169. This paper presents an analysis of {\kep} and {\swift} observations of StHA 169. \begin{figure*} \begin{center} \setlength{\unitlength}{1cm} \begin{picture}(12,12) \put(-3,6.){\special{psfile=allplusq17-llc-mjd-norm.ps hscale=40 vscale=40 angle=0}} \put(-3,-0.4){\special{psfile=all-plusq17-llc-cot-norm-fix-mjd.ps hscale=40 vscale=40 angle=0}} \end{picture} \end{center} \caption{The upper light curve shows the normalised long cadence data of {\src}. The lower light curve has been de-trended and normalised and corrected so that there are no step changes in flux between different quarters.} \label{light} \end{figure*} | The {\sl Swift}-2MASS spectral energy distribution of {\src} indicates there are two components, one relatively hot and one relatively cool. Using additional information such as the existing optical spectrum of Downes \& Keyes (1988) our best estimate for the temperature of the two components is 10000K and 3700K. The {\kep} observations therefore sample the cooler (and physically much larger) component, while the {\swift} UV and blue filters sample the hotter component. The {\kep} light curve shows quasi periodic behaviour with a mean period of 34 d. Given that this period is not stable, it is clearly not the signature of a binary orbital period. Rather it indicates that the cool component is a pulsating red giant star. This is consistent with the suggestion made by Downes \& Keyes (1988) that the cool component in {\src} has a M spectral type and resembles the red giant in the recurrent nova RS Oph (P$_{orb}$=460 d; Dobrzycka \& Kenyon 1994; M0--M2 III, Dobrzycka et al. 1996). Further confirmation of the evolutionary status of the red giant in StHA 169 is provided through its frequency spectrum. {\kep} photometry has been used extensively to characterize red giants as to their membership on the RGB or the AGB (Chaplin et al. 2013). The {\kep} light curve of {\src} is very similar in character to, say, the red giant KIC 2986893 (B{\'a}nyai et al. 2013) which has a mean period of 21.7$\pm$=2.2 d. However, given that {\src} has a wide range of pulsation period, it is possible that it is a Semi Regular Variable (cf Soszy\'{n}ski, Wood \& Udalski 2013). B{\'a}nyai et al. (2013) showed that M giants separate into three distinct groups according to their period structure. StHA 169 and KIC 2986893 belong to group 1 -- red giants with periods between 10--100 days. Group 1 stars with a period similar to that found in {\src} tend to lie the upper Red Giant Branch (see Kiss \& Bedding 2003, 2004) and are pulsating due to first and second overtone modes. (Given the main sequence lifetime of a 3.3 \Msun star is 500 Myr, the system is at least this old). The identification of {\src} as a symbiotic star lies solely with the optical spectrum presented in Downes \& Keyes (1988). The spectral energy distribution as derived from {\swift} and 2MASS photometry and presented in \S \ref{sed} is consistent with {\src} being a binary system. Similarly, our detection of pulsations in the {\kep} data clearly demonstrates that the cool star is a red giant. However, determining the nature of the hot component in symbiotic stars is not a trivial task (see, for instance, Sokoloski \& Bildsten 2010 who recount the quarter of a century debate on the nature of the hot star in the Mira AB system). Our spectral energy distribution shows that an isolated white dwarf or a white dwarf with an accretion disk would not lie at the same infered distances for the red giant component. Instead our results favour that the hot star is more likely to be a late B or early A main sequence star. The absence of short period variability in the UV and the non detection in X-rays suggest that accretion was not taking place at the time of these {\swift} observations. There are at least two other sources which bear some similarity to {\src}: XX Oph and AS 325 which are thought to consist of a Be star and a red giant secondary (Howell, Johnson \& Adamson 2009). Indeed, AS 325 was originally taken to be a symbiotic system. The fact that the optical spectrum of {\src} (Downes \& Keyes 1988) shows the Balmer lines (and He II 4686\AA) in emission may indicate that the B/A star is an emission star (either through a wind or accretion). Stars like these are interesting from a binary evolution point of view. Determining the binary orbital period is a key step but will be difficult to disentangle the signature of the binary period from the red giant pulsations. | 14 | 4 | 1404.5850 |
1404 | 1404.5257_arXiv.txt | { We show that the $\GeV$ scale $\gamma$-ray excess from the direction of the Galactic Center can be naturally explained by the pair annihilation of Abelian vector dark matter (VDM) into a pair of dark Higgs bosons $VV\rightarrow \phi \phi$, followed by the subsequent decay of $\phi$ into $\phi \rightarrow b\bar{b} , \tau \bar{\tau} $. All the processes are described by a renormalizable VDM model with the Higgs portal, which is naturally flavor-dependent. Some parameter space of this scenario can be tested at the near future direct dark matter search experiments such as LUX and XENON1T. } \citestyle{plain} \begin{document} | It has been known for sometime that there is anomalous $\GeV$ scale $\gamma$-ray excess from the direction of the Galactic Center \cite{Goodenough:2009gk,Hooper:2010mq,Boyarsky:2010dr,Hooper:2011ti,Abazajian:2012pn,Gordon:2013vta,Hooper:2013rwa,Huang:2013pda,Abazajian:2014fta,Daylan:2014rsa}. Interestingly, the excess seems to be remarkably well described by an expected signal from 31-40 $\GeV$ dark matter (DM) annihilating dominantly to $b\bar{b}$ with a cross section $\sigma v \simeq \l(1.7-2.3\r) \times 10^{-26} {\rm cm}^3/{\rm s}$ \cite{Daylan:2014rsa,Steigman:2012nb}. (See also Ref.~\cite{Yuan:2014rca} for millisecond pulsars as an astrophysical alternative.) Because of the importance of DM pair annihilation into $b\bar{b}$ for the GC $\gamma$-ray excess, some ideas on flavored DM have been put forth~\cite{Berlin:2014tja,Agrawal:2014una}. We should note that it is the shape of $\gamma$ spectrum from dark matter annihilation that mainly matters rather than the precise value for $\sigma v$ since there is a large uncertainty in the density profile of dark matter near the Milky Way center. As long as $\langle \sigma v\rangle \left( \rho_{\textrm{DM}}/M_\textrm{DM} \right)^2$ is at the right amount, a good fit can be achieved for $b\bar{b}$ final state. Actually, $b\bar{b}$ does not need to be the only annihilation channel, it was shown~\cite{Daylan:2014rsa} that flavor-dependent annihilations can also fit the data well. Such kind of flavor-dependent annihilations may indicate a Higgs-like scalar mediator, since Higgs-like scalar will couple with the heaviest particle it can couple to. The required cross section is very close to the canonical value for neutral thermal relic dark matter. It can be achieved either $s$-wave annihilation or $p$-wave annihilation with $s$-channel resonance at present. However, in the latter case, the resonance band is likely to be very narrow that leads to a severe fine-tuning, which is not that attractive. With this consideration, perhaps the simplest scenario for dark matter model that can explain the $\gamma$-ray excess would be those involving scalar mediator with Higgs portal interaction(s), since in this case the scalar mediator will couple strongly to the $b\bar{b}$, the heaviest particles kinematically producible \footnote{Another possibility would be to consider flavored DM ~\cite{Berlin:2014tja,Agrawal:2014una}. }. Then, one can imagine the following simple scenarios of DM having $s$-wave annihilation channel: \begin{enumerate} \item Singlet scalar dark matter (SSDM): a real scalar mediator \cite{Modak:2013jya} \item Singlet fermion dark matter (SFDM): a pseudo-scalar mediator \cite{Boehm:2014hva,Izaguirre:2014vva,Ipek:2014gua} \footnote{While we were working on these possibilities, this paper was put on the archive, and we don't consider this possibility any more in this work.} \item Singlet vector dark matter (SVDM): a real scalar mediator \cite{Boehm:2014bia,Baek:2012se,Baek:2014goa} \end{enumerate} Note that the structure of above scenarios can be realized easily when DM is charged under a dark gauge symmetry which is broken to, for example, a discrete $Z_2$ or $Z_3$ symmetry. Hence those scenarios would also work equally well. For other recent proposals of DM models to address the $\GeV$ $\gamma$-ray spectrum, see Refs. \cite{Okada:2013bna,Alves:2014yha,Berlin:2014tja,Agrawal:2014una}. Potentially the most important constraint on those singlet dark matter models may come from direct search experiments, for example, LUX \cite{Akerib:2013tjd}. However the existence of extra scalar boson mediating dark and visible sectors via Higgs portal interaction(s) has a significant effect on direct searches if the mass of the extra non-SM Higgs is not very different from that of SM Higgs \cite{Baek:2011aa}, and the constraint from direct searches can be satisfied rather easily. Note that this feature is not captured at all in effective field theory approach, and it is important to work on the minimal renormalizable and unitary Lagrangian for physically sensible results \footnote{See Refs.~\cite{Baek:2011aa,Baek:2012se} for the original discussions on this point, and Ref.~\cite{HiggsPortalEFT} for more discussion on the correlation between the invisible Higgs branching ratio and the direct detection cross section in the Higgs portal SFDM and SVDM models,}. In this paper, we revisit SVDM scenario with Higgs portal in the context of the the $\gamma$-ray excess from the Galactic Center, and show that the SVDM model can naturally explain it, while satisfying all of known constraints coming from CMB, Fermi-LAT $\gamma$-ray search and LHC experiments. We also show that the parameter space relevant for the $\gamma$-ray excess can be probed by the near future direct dark matter search experiment, for example LUX and XENON1T. This paper is organized as follows. In Section~\ref{sec:model}, we recapitulate the renormalizable SVDM model with Higgs portal. In Section~\ref{sec:const}, various relevant constraints on the model are discussed, including relic density estimation, vacuum stability, collider bounds, CMB and direct detection cross section, etc., and we show that our model can explain the $\gamma$-ray excess from the galactic center without any conflict with other cosmological and astrophysical observations. In Section~\ref{sec:conc}, our conclusion is drawn. | \label{sec:conc} In this paper, we revisited the singlet vector dark matter (SVDM) model with Higgs portal in order to see if it can explain the observed $\GeV$ scale $\gamma$-ray spectrum from the galactic center by the annihilation of dark matter mainly to $b\bar{b}$ or to two non-SM light Higgses which decay subsequently and dominantly to $b\bar{b}$. We found that the Higgs portal SVDM scenario can naturally explain the $\gamma$-ray spectrum while providing a right amount of relic density for $m_h/2 < m_V \lesssim 80 \GeV$ and $(m_V-m_\phi)/m_V \ll 1$ with $m_V$ and $m_\phi$ being the masses of VDM and non-SM Higgs boson. This implies that the mass of the non-SM extra Higgs is constrained to be within a narrow range of \beq m_h/2 \lesssim m_\phi \lesssim 80 \GeV \eeq which can be focused on in dedicated searches of the second Higgs at future collider experiments although a null result due to very small mixing angle $\alpha$ is also possible. The dark gauge coupling is contained to be $g_X \sim 0.2$ for the right amount of relic density while taking $\alpha$ to be small enough to satisfy direct DM search bound. Unfortuantely the LUX or XENON1T cannot explore the entire parameter space of the SVDM explaining the GeV-scale $\gamma$-ray from the galactic center. The instability of SM vacuum could be improved due to the additional loop contribution of an extra scalar field. | 14 | 4 | 1404.5257 |
1404 | 1404.0228_arXiv.txt | This essay addresses the issue of gravitational phase transitions in the early universe. We suggest that a second order phase transition observed in the Causal Dynamical Triangulations approach to quantum gravity may have a cosmological relevance. The phase transition interpolates between a non-geometric crumpled phase of gravity, and an extended phase with classical properties. Transitions of this kind have been postulated earlier in the context of geometrogenesis in Quantum Graphity. We show that critical behavior may also be associated with a signature change event in Loop Quantum Cosmology. In both cases, classical spacetime originates at the critical point associated with a second order phase transition. | Accumulating results of theoretical investigations indicate that the gravitational field exists in different phases. First indications supporting such an idea came from considerations of three-dimensional Euclidean quantum gravity \cite{Ambjorn:1991rx}. By means of Monte Carlo simulations it was possible to explore the configuration of the gravitational field under various conditions. For four-dimensional Euclidean gravity, gravity exhibits two phases: the \emph{crumpled} phase and the \emph{branched polymer} phase \cite{Ambjorn:1991pq}. This result has since been generalized to the case of four-dimensional gravity with an imposed causality condition, formulation known as Causal Dynamical Triangulations (CDT). The causality condition turned out to be essential for the correct phase structure of gravity, leading to emergence of the four-dimensional spacetime \cite{Ambjorn:2004qm}. Generation of such extended phase in the Euclidean approach introducing a non-trivial path integral measure remains an interesting possibility \cite{Laiho:2011ya}. Furthermore, in similarity to a phase structure of the Lifshitz scalar \cite{Hornreich:1975zz}, the critical surface of CDT has been divided into three regions separated by the first and second order transition lines \cite{Ambjorn:2012ij}. Interestingly, a theory describing gravity at a triple point (Lifshitz point) of the phase diagram has been constructed, and shown to be power-counting renormalizable \cite{Horava:2009uw}. Further evidence for the non-trivial phase structure of gravity comes from Quantum Graphity \cite{Konopka:2006hu}. This approach utilizes the idea of \emph{geometrogenesis} - a transition between geometric and non-geometric phases of gravity. The basic question one can ask, assuming the existence of the different phases of gravity, is where the other phases can be found? A natural place to search for them are high curvature regions such as interiors of the black holes and the early universe. Because of a horizon, a possibility of empirical verification of a phase change inside of black holes is out of reach. More promising is a search for signatures of the gravitational phase transitions which took place in the early universe. So far, there has been very little attention devoted to this issue in the literature. Most studies of the phase transitions in the early universe were dedicated to the matter sector, rather than gravity \cite{Kibble:1976sj,Zurek:1985qw}. Among the few studies on the gravitational phase transitions in the early universe, the work of Refs. \cite{Magueijo:2006fu,Dreyer:2013vka} is especially noteworthy. In Ref. \cite{Magueijo:2006fu} a specific model of geometrogenesis, through a second order phase transition, has been proposed. It was shown that by assuming the holographic principle to be fulfilled in the high temperature phase, it is possible to generate a power spectrum of primordial perturbations that is in agreement with observations. In Ref. \cite{Dreyer:2013vka} the cosmological relevance of second order phase transitions is discussed. Arguments supporting generation of ``inflationary'' power spectrum from critical behavior of the gravitational field have been presented. In what follows we attract attention to the fact that a second order gravitational phase transition has recently been observed within Causal Dynamical Triangulations \cite{Ambjorn:2011cg}. The transition takes place exactly between the phases of the form discussed in Refs. \cite{Magueijo:2006fu,Konopka:2006hu}. Therefore, CDT gives a concrete realization of the scenario of geometrogenesis. We also show that gravitational phase transition may be associated with the deformation of general covariance, recently observed in the context of Loop Quantum Cosmology (LQC). In both cases, the phase transition is of second order, suggesting a critical nature of the emergence of classical spacetime in the early universe. | We have shown that CDT offers a concrete realization of geometrogenesis, having a second order gravitational phase transition between the non-geometric and geometric phase. We have explained how measurements of the spectral dimension are related in this case with connectivity of the chunks of space. The critical nature of the emergence of classical spacetime in the early universe may give a first possibility of testing CDT. However, this would require more detailed investigations of the properties of the second order phase transition in CDT. In LQC, which is an alternative to CDT, a gravitational second order phase transition may explain a signature change in the Planck epoch. In the presented toy model, the universe originates just at the critical point. Interestingly, the latest results in CDT indicate the existence of a new phase between phases B and C \cite{Ambjorn:2014mra}, characterized by a ``bifurcation'' of the kinetic term. This behavior resembles the signature change observed in LQC. | 14 | 4 | 1404.0228 |
1404 | 1404.5311_arXiv.txt | We investigate the collapse of primordial gas in a minihalo with three-dimensional radiation hydrodynamics simulations that accurately model the transfer of H$_2$ line emission. For this purpose, we have implemented a multiline, multifrequency ray-tracing scheme in the moving-mesh code {\sc arepo} that is capable of adaptively refining rays based on the {\sc healpix} algorithm, as well as a hybrid equilibrium/non-equilibrium primordial chemistry solver. We find that a multifrequency treatment of the individual H$_2$ lines is essential, since for high optical depths the smaller cross-section in the wings of the lines greatly increases the amount of energy that can escape. The influence of Doppler shifts due to bulk velocities is comparatively small, since systematic velocity differences in the cloud are typically smaller than the sound speed. During the initial collapse phase, the radially averaged escape fraction agrees relatively well with the fit of Ripamonti \& Abel. However, in general it is not advisable to use a simple density-dependent fitting function, since the escape fraction depends on many factors and does not capture the suppression of density perturbations due to the diffusion of radiation. The Sobolev method overestimates the escape fraction by more than an order of magnitude, since the properties of the gas change on scales smaller than the Sobolev length. | In the standard $\Lambda$ cold dark matter ($\Lambda$CDM) cosmology, the first stars are expected to form at redshifts $z\ga 20$ in dark matter (DM) `minihaloes' with masses $\sim 10^6\,{\rm M_\odot}$ \citep[for recent reviews, see][]{bromm13, glover13}. The primordial gas synthesized in the big bang accretes on to these haloes and heats to the virial temperature of $T_{\rm vir}\simeq 1000\,{\rm K}$. At the centre of the halo, molecular hydrogen (H$_2$) forms via associative detachment of neutral hydrogen with negatively charged hydrogen (H$^-$), which in turn forms via radiative association of neutral hydrogen with free electrons left over after recombination \citep{sz67}. The internal ro-vibrational transitions of H$_2$ are excited by collisions with other species, and their decay produce cooling radiation that facilitates the further collapse of the gas. Detailed three-dimensional simulations of primordial star formation have shown that the onset of H$_2$ cooling in sufficiently massive haloes results in runaway collapse to a density of $n_{\rm H}\simeq n_{\rm cr}=10^4\,{\rm cm}^{-3}$, where $n_{\rm H}$ denotes the volumetric number density of hydrogen nuclei, and $n_{\rm cr}$ is the critical density at which the ro-vibrational levels of H$_2$ become populated according to local thermal equilibrium \citep[LTE;][]{abel98, bcl99, bcl02, abn00, abn02}. The gas cools to a minimum temperature of $\simeq 200\,{\rm K}$, which is set by the change in the scaling of the cooling rate from $\Lambda\propto n_{\rm H}^2$ to $\Lambda\propto n_{\rm H}$ for $n_{\rm H}\ga n_{\rm cr}$, and the microphysics of the H$_2$ molecule. The collapse rate decreases and the cloud `loiters' at this density and temperature until it has accreted a Jeans mass of $\ga 100\,{\rm M_\odot}$. The cloud then decouples from the DM potential and begins to collapse under its own gravity. At densities $n_{\rm H}\ga 10^8\,{\rm cm}^{-3}$, the H$_2$ fraction rapidly increases due to three-body reactions \citep{pss83}, and the rapidly increasing cooling rate may trigger a chemothermal instability that results in subfragmentation of the cloud \citep{sy77, silk83, ra04, yoshida06b, tao09, gsb13}. At densities $n_{\rm H}\ga 10^{10}\,{\rm cm}^{-3}$, the stability of the cloud is restored due to the increasing optical depth of the gas to H$_2$ line emission \citep{on98, ripamonti02, ra04, yoshida06b}. The second radiative coolant that becomes important is collision-induced emission, which operates at densities $n_{\rm H}\ga 10^{14}\,{\rm cm}^{-3}$ \citep{on98, ra04}. Finally, chemical cooling due to the dissociation of H$_2$ molecules precedes the near-adiabatic evolution of the gas at the highest densities \citep{on98, yoh08}. One of the most important transitions that occurs during the initial collapse phase is the transition from optically thin to optically thick H$_2$ line cooling. Previous studies have shown that the chemical and thermal evolution of the gas at high densities depend sensitively on how rapid this transition occurs \citep{turk11, hy13}. However, an accurate solution requires the hydrodynamic evolution of the gas to be solved alongside the multifrequency radiative transfer of H$_2$ line emission. This has so far only been possible in one-dimensional calculations \citep{on98, ripamonti02}. In three-dimensional simulations, the computational cost associated with the integration of the six-dimensional photon distribution function is prohibitively expensive. Previous studies have therefore resorted to the escape probability formalism, where the optically thin cooling rate is multiplied by an escape fraction that models the probability of a photon to escape from the cloud. The escape fraction is usually derived from local properties of the gas. In \citet{ra04}, a density-dependent fit for the escape fraction was obtained from the detailed one-dimensional calculations of \citet{ripamonti02}. This method was used in \citet{tao09, tna10, turk12}. Other studies assumed that the velocity gradient in the central, Jeans-unstable cloud allows the radiation to escape relatively easily, and used the \citet{sobolev60} method to obtain the escape fraction \citep[e.g.][]{yoshida06b, clark11a, gsb13}. The Sobolev method has been traditionally applied to stellar atmospheres and molecular clouds \citep{castor70, gk74}, and is valid if the scale on which the velocity varies is much smaller than the scale on which other properties of the gas vary. In particular, it is not well suited to treat turbulent gas clouds \citep{schoenberg85, ossenkopf97}. Nevertheless, the results obtained with the Sobolev method in three-dimensional simulations agree relatively well with those of the fitting function of \citet{ra04}, even though is not clear how accurate both methods really are \citep{turk11, hy13}. We here address this issue by performing the first three-dimensional simulations of primordial star formation that include multifrequency radiative transfer for optically thick H$_2$ line emission. In Section~\ref{sec_chem}, we present a new primordial chemistry and cooling network implemented in {\sc arepo}, and in Section~\ref{sec_rad} describe the ray-tracing scheme used to compute the radiative transfer. In Section~\ref{sec_sim}, we describe the set-up of the simulations, and in Section~\ref{sec_res} present our results. Finally, in Section~\ref{sec_sum} we summarize and draw conclusions. All distances are quoted in proper units, unless noted otherwise. | \label{sec_sum} We have performed the first three-dimensional simulations of primordial star formation that self-consistently model the multifrequency radiative transfer of H$_2$ line emission. The simulations employ a new equilibrium/non-equilibrium primordial chemistry solver next to a new multiline, multifrequency ray-tracing scheme that is capable of adaptively refining rays based on the {\sc healpix} algorithm \citep{gorski05}. The latter can be used to solve the static radiative transfer equation for point sources as well as diffuse emission. Both schemes have been implemented in the simulation code {\sc arepo}. The chemistry solver is optimized for collapse simulations and is significantly faster than the solver used in \citet{greif12}. The ray-tracing scheme is capable of walking about one million cells per second, and can be parallelized using a hybrid distributed/shared memory scheme. The calculation of the optical depth for multiple lines and/or frequency bins uses tabulated opacities and exploits the Intel AVX instruction set, which boosts the performance to a level where the main bottleneck is the communication of the ray data among {\sc MPI} tasks. For isolated point sources, the parameters that govern the accuracy of the scheme are the initial {\sc healpix} level, the average number of rays per cell (for adaptive splitting), the minimum fractional energy of a ray before it is terminated, the number of lines, and the number of frequency bins. In the case of diffuse emission, the angular resolution typically remains constant, while an additional parameter specifies the fraction of all cells that are sources. The reliability of both schemes is demonstrated with a series of idealized test calculations. The ray-tracing scheme is used to compute the radiative transfer of H$_2$ line emission in an ab initio simulation of primordial star formation. We achieve an accuracy of $5$ per cent in the radiative heating rate by using $\simeq 10^{14}$ opacity calculations per time step, amounting to a total wall-clock time of $1$--$2$ months on $1024$ state-of-the-art computing cores. In agreement with previous studies, we find that the gas becomes optically thick to H$_2$ line emission at densities $n_{\rm H}\ga 10^{10}\,{\rm cm}^{-3}$, and the line cooling rate is surpassed by collision-induced emission at $n_{\rm H}\ga 10^{15}\,{\rm cm}^{-3}$. Within this range, the spherically averaged escape fraction decreases from unity to $\simeq 2\times 10^{-3}$, with a power-law slope of $\simeq -0.6$. This agrees relatively well with the fitting function of \citet{ra04}, which is based on one-dimensional radiative transfer calculations \citep{on98, ripamonti02}. During the initial collapse phase, the assumption of spherical symmetry appears to give relatively accurate results for the purpose of computing the H$_2$ line transfer. However, since the escape fraction depends on many factors, such as the density, velocity, temperature, and the chemical and thermal rate equations, it is generally not advisable to use a fitting function. The escape fraction method also does not capture the diffusion of the radiation, which suppresses density fluctuations as the gas evolves into the optically thick regime. By systematically increasing the physical detail of the radiative transfer, we have found that using multiple lines and frequency bins is essential. The lower cross-sections of the more sparsely populated lines can boost the amount of energy that can escape by many orders of magnitude. A similar effect becomes important if frequency-dependent emission and absorption are accounted for: the lower cross-sections in the wings of the lines allow significantly more energy to escape than in the grey case. Finally, Doppler shifts due to relative velocities along the rays increase the escape fraction by about a factor of $2$. This effect is relatively small in comparison, since the infall velocity fluctuates by less than the sound speed, and results in a frequency shift of only $\simeq \Delta\nu_{\rm D}/2$. We have also compared our results to the escape fraction obtained with the Sobolev method. For low optical depths, the Sobolev method somewhat underestimates the escape fraction, while for high optical depths the escape fraction is overestimated by more than an order of magnitude. This discrepancy arises because the Sobolev method is only accurate if the scales on which the properties of the gas change are much larger than the Sobolev length. This is not the case in the self-gravitating gas clouds that form in minihaloes, since the infall velocity typically varies by less than the sound speed. The discrepancy becomes even larger if the turbulence is well resolved. In this case, the local velocity gradient can be much larger than $c_{\rm s}/\lambda_{\rm J}$, resulting in a further increase in the escape fraction, despite the fact that the turbulent velocities nearly cancel each other within a Jeans length, and thus have almost no effect on the escape fraction. Previous studies found better agreement between the Sobolev method and the fitting function of \citet{ra04}, since limitations of the hydrodynamic solver employed prevented the turbulence from being resolved \citep{yoshida06b, clark11b, turk11, hy13}. As a result, the velocity gradient was dominated by the radial velocity gradient, resulting in a Sobolev length that was significantly larger than the one found here. For the above reasons, simulations that used the Sobolev method and resolved the turbulence in the gas greatly overestimated the escape fraction \citep{greif12, gsb13}. In particular, the cooling instability found in \citet{gsb13} is largely an artefact of the Sobolev method. However, due to the strong dependence of the H$_2$ line cooling rate on the temperature, the overall thermal evolution of the cloud is much less affected. For example, in \citet{gsb13} the temperature at $n_{\rm H}\simeq 10^{15}$ is $\simeq 1800\,{\rm K}$, while in \citet{tao09} the temperature is $\simeq 2500\,{\rm K}$. The H$_2$ abundance shows a similarly mild variation. In \citet{gsb13}, the gas remains fully molecular at a density of $\simeq 10^{15}$, while in \citet{tao09} the H$_2$ has begun to dissociate, with $y_{\rm H_2}\simeq 0.2$. \citet{turk12} suggested that the reduced H$_2$ fraction may reduce the ability of the cloud to fragment. However, since the gas becomes rotationally supported in a Keplerian disc following the initial collapse, the resulting asymmetry may allow the cooling radiation to escape more easily than previous studies predicted \citep{clark11a, greif11a, greif12}. A definitive answer must await detailed radiation hydrodynamics simulations that evolve the collapse well beyond the formation of the first protostar. | 14 | 4 | 1404.5311 |
1404 | 1404.3701_arXiv.txt | { Observations of Faraday rotation for extragalactic sources probe magnetic fields both inside and outside the Milky Way. Building on our earlier estimate of the Galactic contribution, we set out to estimate the extragalactic contributions. We discuss the problems involved; in particular, we point out that taking the difference between the observed values and the Galactic foreground reconstruction is not a good estimate for the extragalactic contributions. We point out a degeneracy between the contributions to the observed values due to extragalactic magnetic fields and observational noise and comment on the dangers of over-interpreting an estimate without taking into account its uncertainty information. To overcome these difficulties, we develop an extended reconstruction algorithm based on the assumption that the observational uncertainties are accurately described for a subset of the data, which can overcome the degeneracy with the extragalactic contributions. We present a probabilistic derivation of the algorithm and demonstrate its performance using a simulation, yielding a high quality reconstruction of the Galactic Faraday rotation foreground, a precise estimate of the typical extragalactic contribution, and a well-defined probabilistic description of the extragalactic contribution for each data point. We then apply this reconstruction technique to a catalog of Faraday rotation observations for extragalactic sources. The analysis is done for several different scenarios, for which we consider the error bars of different subsets of the data to accurately describe the observational uncertainties. By comparing the results, we argue that a split that singles out only data near the Galactic poles is the most robust approach. We find that the dispersion of extragalactic contributions to observed Faraday depths is most likely lower than $\SI{7}{\radian/m^2}$, in agreement with earlier results, and that the extragalactic contribution to an individual data point is poorly constrained by the data in most cases.} | Polarized radiation from an astronomical source undergoes Faraday rotation as it travels through the magneto-ionic medium between the source and observer. For extragalactic sources, there are contributions from the Galactic interstellar medium, from any intergalactic magnetic fields, from intervening galaxies on the line of sight, as well as from magnetic fields in the source itself. In this work we attempt to estimate the contribution to the observed Faraday rotation of such sources that is due to magnetic fields outside of the Milky Way. This extragalactic contribution holds the potential for extracting information about cosmic magnetic fields on large scales, e.g., in galaxy clusters, galaxy filaments, or cosmic voids \citep{kolatt-1998, blasi-1999, xu-2006, hammond-2012, bernet-2012, neronov-2013, joshi-2013}. For a hypothetical source of linear polarization that is point-like in all dimensions and situated at a physical distance $r$ from the observer, the change in polarization angle is given by \begin{equation} \Delta_\chi = \phi \, \lambda^2, \end{equation} where $\lambda$ is the wavelength of the radiation and \begin{equation} \label{eq:def_phi} \phi = \frac{\mathrm{e}^3}{2\pi \, m_\mathrm{e}^2 \, c^4} \int_r^0 \mathrm{d}r' \, \frac{1}{\left(1 + z(r')\right)^2} \, n_\mathrm{e}{(r')} \, B_r{(r')} \end{equation} is the Faraday depth of the source \citep[e.g.,][]{burn-1966}. In the last equation, $n_\mathrm{e}$ is the density of thermal electrons, $B_r$ is the magnetic field vector projected onto the line of sight, $z$ is the cosmological redshift, and the prefactor is a function of the electron charge $\mathrm{e}$, the electron mass $m_\mathrm{e}$, and the speed of light $c$. The line-of-sight integral in Eq.~\eqref{eq:def_phi} can be split into an integral over the portion of the line of sight that lies within the Milky Way and the portion that is outside the Milky Way, i.e., the Galactic and extragalactic contributions. For most extragalactic sources, the net rotation is dominated by the effect of the interstellar medium of the Milky Way \citep{leahy-1987, schnitzeler-2010}. This Galactic contribution has recently been estimated from a collection of observations of Faraday rotation of extragalactic sources \citep{oppermann-2012}. One way of estimating the extragalactic contributions is to subtract the estimate of the Galactic contribution from the observed values. However, we will argue that this is not a good estimate due to the presence of uncertainties both in the observations and in the foreground estimate. Extracting the sub-dominant extragalactic part from the data is more difficult than extracting the Galactic part for two reasons. The first obvious reason is that, as it is smaller, it is more heavily obscured by observational noise in the data. In fact, for many of the data points that we use the expected extragalactic contribution is comparable to the observational uncertainty. The second reason is that the Galactic foreground contribution is spatially smooth to some extent, which enables the usage of correlation information and thus the transfer of information from many data points to each location on the sky. The extragalactic contributions, on the other hand, are expected to be mostly uncorrelated from source to source, given the typical source separation. Therefore, information on the extragalactic contribution to a data point is only contained in the other data points indirectly via the common Galactic foreground. Furthermore, the measurement errors are uncorrelated from source to source as well leading to a statistical degeneracy with the extragalactic contributions. This means that any split between extragalactic contributions and observational noise in the estimate can only be made according to the expected variances of these two components. We will additionally argue that the statistical characterization of the observational uncertainties given by the error bars in the data catalogs may be incomplete in some cases. Therefore, there is an additional degree of freedom in the expected noise variance that further exacerbates the degeneracy between observational noise and extragalactic contributions. The dispersion of the extragalactic contributions has previously been estimated by \citet{schnitzeler-2010}, who studied the spread of the distribution of observed Faraday depths of extragalactic sources from the catalog of \citet{taylor-2009}. \citet{schnitzeler-2010} observed that this spread changes as a function of Galactic latitude even after the subtraction of a coarse foreground model. He then extracted an upper bound on the spread of the extragalactic contributions as the latitude independent part of this function. Here we will regard this spread as unknown and derive a complementary estimate. For this, we will separate the data into a subset for which the uncertainty information is complete and reliable and a subset for which this is not necessarily the case. The reconstruction of the dispersion of the extragalactic contributions will then be driven mainly by the former subset of the data. We will argue that the best choice for the first subset consists of data that have not only well-described observational uncertainties, but also a small Galactic foreground contribution, i.e., we will prefer data from the Galactic polar regions. The Galactic contribution will be separated off by considering its different spatial correlation structure. In the following Sect.~\ref{sec:terminology}, we give precise definitions for the terminology that we use in the later discussion. Terms like noise and estimate are introduced and we discuss the relevant probability densities. In Sect.~\ref{sec:algorithm}, we sketch the derivation of the reconstruction algorithm and discuss the degeneracy between the extragalactic and noise contributions. We test the algorithm on a simulation. In discussing the resulting estimates for the extragalactic contributions, we point out the important difference between an estimate and the reality, as well as the importance of considering estimates together with their uncertainties. In Sect.~\ref{sec:realworld}, we apply the algorithm to observational data and present the results. We perform different case studies to gauge the robustness of these results. Finally, we give a brief summary in Sect.~\ref{sec:conclusions}. A reader interested mainly in the results may skip the derivation and test of the method in Sect.~\ref{sec:algorithm} and go straight from Sect.~\ref{sec:terminology} to Sect.~\ref{sec:polarcapsresults}. In that section, we discuss the results for the split of the data described as the second split in Sect.~\ref{sec:possiblesplits}, which we argue gives the most reliable results. A discussion of the use of these results is given in Appendix~\ref{app:website}. | \label{sec:conclusions} We have studied the contributions to the observed Faraday rotation of extragalactic sources that are due to the Galactic interstellar medium, due to extragalactic magnetic fields, and due to observational noise. Extracting any of these three contributions is non-trivial, as they are superimposed on every line of sight. Another complication is that the observational error bars do not in every case describe the data likelihood accurately. This makes even a probabilistic analysis of the fractions of the data values due to the three different constituents challenging. If the observations were noiseless, the extragalactic contributions could be estimated by simply subtracting an estimate of the Galactic foreground from the data values. However, in reality the observations are noisy and an estimate of the extragalactic contributions calculated in this way will contain this noise as well. Simply subtracting a Galactic foreground from the data is therefore not a good way of estimating extragalactic contributions. Furthermore, any estimate of the Galactic foreground will itself be uncertain and this uncertainty, when not taken properly into account, will introduce artifacts in the extragalactic estimate. In our considerations, we strictly made the distinction between a physical quantity and an estimate of this quantity. The latter aims to equal the former, but, even if calculated correctly, there is always uncertainty involved and artifacts in the estimate may result. Taking into account the uncertainty of the estimate, however, should remove the artifacts. An example of such an artifact is the latitude dependence that we observed in the posterior mean estimates for the extragalactic Faraday depths. This latitude dependence vanishes once the uncertainty is taken into account. To treat the complete problem of estimating the amount of both Galactic and extragalactic Faraday rotation from observations, we extended the algorithm of \citet{oppermann-2012}. This extended algorithm is based on a split of the data into a subset for which the observational error bars describe the data likelihood sufficiently and another subset for which this is not the case. It includes the estimation of the angular power spectrum of the Galactic foreground, assumed to be statistically isotropic up to a single latitude-dependent modulation, the estimation of this latitude-dependent function, the estimation of corrected noise variances for the subset of the data for which this is deemed necessary, and the estimation of the variance of the extragalactic contributions. We showed in a simulated scenario that all of these quantities are accurately reconstructed by our algorithm if our statistical model, including the split of the data, is correct. For the application to observational data, we have considered several different ways to split the data into the two categories. We find that the most robust outcomes are achieved with splits that only regard a small fraction of the data (we use $1.75\%$ of the data points) situated near the Galactic poles as not afflicted by potential problems in the description of the data likelihood. In these cases we find extragalactic dispersions between $\SI{6.6}{\radian/m^2}$ and $\SI{7.2}{\radian/m^2}$. These numbers agree remarkably with the ones derived by \citet{schnitzeler-2010} by splitting the dispersion of observed Faraday rotation values into a latitude-dependent part, a contribution due to measurement errors, and a constant offset, deemed to be extragalactic in origin. Strictly speaking, both analyses only produce upper limits on the dispersion of the extragalactic contributions, but for slightly different reasons. While the estimate of \citet{schnitzeler-2010} may be increased due to a latitude-independent Galactic contribution, our estimate may be increased due to a Galactic contribution that is spatially uncorrelated on the scales probed by the observations. We provide the derived estimates for all the involved quantities online at \url{http://www.mpa-garching.mpg.de/ift/faraday/}. We explain in Appendix~\ref{app:website} how these estimates can be used to estimate related quantities. The foreground products can be seen as updated versions of the results of \citet{oppermann-2012} that should be used preferentially, except in special cases where one of the assumptions we made in this paper is at question. We also provide 1\,000 samples of extragalactic contributions to the observed Faraday rotation, drawn from the posterior PDF for this quantity. This will enable future studies of extragalactic Faraday rotation to take into account the full probability distribution for these values, by performing any analysis on the set of samples rather than only on the posterior mean estimate. It should be noted that, within the framework of our assumptions, the extragalactic contributions are not very well constrained by the data. This is to some extent due to allowing the observational error bars of sources to get increased during the reconstruction, which increases the uncertainty of all reconstructed quantities. In addition, sources for which such an increase of the error bar can happen in our reconstruction will not have large estimates of the extragalactic contribution. All our considerations point toward the importance of understanding the uncertainties of Faraday rotation measurements. For future surveys, this means that not only should the largest possible interval in $\lambda^2$-space be covered, but, as already pointed out by \citet{farnsworth-2011} and \citet{farnes-2014}, all the available information should be used in the data reduction, including the behavior of polarization fraction with frequency, as this can help avoid some of the rather poorly understood effects in RM synthesis studies that can lead to faulty estimates. | 14 | 4 | 1404.3701 |
1404 | 1404.1018_arXiv.txt | { This article lays out a complete framework for an effective theory of cosmological perturbations with corrections from canonical quantum gravity. Since several examples exist for quantum-gravity effects that change the structure of space-time, the classical perturbative treatment must be rethought carefully. The present discussion provides a unified picture of several previous works, together with new treatments of higher-order perturbations and the specification of initial states.} | One of the most promising avenues towards tests of quantum gravity is the study of the corrections it implies in cosmological perturbation equations. With some luck, characteristic traces of high-energy space-time phenomena may be left in the observable part of the cosmic microwave background. The derivation of power spectra from different approaches to quantum gravity is therefore one of the most active areas in the field. In this article, we consider the status of developments in canonical quantum gravity, in particular loop quantum gravity. Canonical quantum gravity does not straightforwardly lead to cosmological perturbation equations. Given a consistent quantum theory, one would have to find a suitable class of semi-classical states, so that the expectation values they define for basic operators (such as the spatial metric or its inhomogeneous modes) are subject to equations of motion derived from the Wheeler--DeWitt equation. The issue of finding semi-classical (or other suitable) states in canonical quantum gravity is tricky --- for one, we are unaware of any ground state that would be a good starting point for a perturbative expansion. And if we assume that the problem can be solved, consistent cosmological perturbation equations do not follow automatically, for reasons closely related to the generally covariant nature and complicated gauge content of the theory. The classical theory has more equations than unknowns, and they appear as different types. There are constraints that depend only on the modes and their first time derivatives (or on configuration variables and their canonically conjugate momenta), and there are equations of motion that determine second-order time derivatives. The classical set of equations satisfies two important consistency conditions: (i) The constraints, a spatial function $C$ and a vector field $\vec{D}$, are preserved by the equations of motion, so that they need to be imposed only for initial values and then automatically hold at all later times. This preservation of the constraints is the reason why the set of equations for a smaller number of free functions is consistent and allows the right number of non-trivial solutions. (ii) The constraints are generators of a large set of gauge transformations which include space-time Lie derivatives of phase-space functions, in particular their time derivatives. Time derivatives of momentum variables imply the second-order equations of motion; by being generated by gauge-inducing constraints, they automatically preserve the constraints. In canonical language, the equations of motion are generated by the constraints, that is they follow from equations $\dot{f}=\{f,H[N,\vec{M}]\}$ in which the time derivative (along a direction with space-time components related to the lapse function $N$ and the shift vector $\vec{M}$) of a mode function $f$ is given by the Poisson bracket with a Hamiltonian $H[N,\vec{M}]$. This Hamiltonian is a linear combination $H[N,\vec{M}]=\int(NC+\vec{M}\cdot\vec{D})$ of the constraint functions $C$ and $\vec{D}$ and therefore must vanish when evaluated for solutions of the field equations, for all $N$ and $\vec{M}$. Property (i) is guaranteed if $0=\dot{H}[N_1,\vec{M}_1]= \{H[N_1,\vec{M}_1], H[N,\vec{M}]\}$ for all $N_1$ and $\vec{M}_1$ and for all configurations for which $H[N_1,\vec{M}_1]=0$, in which case the constraints are called first-class. The gauge transformations mentioned in property (ii) take the form $\delta_{(\epsilon_0,\vec{\epsilon})}f= \{f,H[\epsilon_0,\vec{\epsilon}]\}$. For the specific choice of $(\epsilon_0,\vec{\epsilon})=(N,\vec{M})$, one obtains time derivatives as gauge transformations. The general gauge transformation with parameters $(\epsilon_0,\vec{\epsilon})$ amounts to a Lie derivative along a space-time vector field with components $(\epsilon_0/N, \vec{\epsilon}- \epsilon_0\vec{M}/N)$. Property (ii) then has the following consequence: By choosing different linear combinations, varying $N$ and $\vec{M}$ in the Hamiltonian, one takes into account all possible time choices. In other words, the theory is invariant under reparametrizations of time. It is therefore possible to express the equations of motion in terms of only gauge-invariant (i.e., coordinate independent) combinations of the modes. No gauge artifacts couple to the physical degrees of freedom. These consistency conditions are a specific realization of the general concept of a first-class system of constraints, using the notation introduced by Dirac. There is a set of constraints, in our case $H[N,\vec{M}]$, or individually $C[N]:=\int NC=0$ and $\vec{D}[\vec{M}]:=\int \vec{M}\cdot\vec{D}=0$, so that their Poisson brackets $\{\vec{D}[\vec{M}_1],\vec{D}[\vec{M}_2]\}$, $\{C[N],\vec{D}[\vec{M}]\}$ and $\{C[N_1],C[N_2]\}$ vanish whenever the constraints are satisfied. In the case of gravity, the constraints obey the hypersurface-deformation algebra \cite{DiracHamGR} \begin{eqnarray} \{\vec{D}[\vec{M}_1],\vec{D}[\vec{M}_2]\} &=& -\vec{D}[{\cal L}_{\vec{M}_2}\vec{M}_1]\,,\label{DD}\\ \{C[N],\vec{D}[\vec{M}]\} &=& -C[{\cal L}_{\vec{M}}N]\label{HD}\,,\\ \{C[N_1],C[N_2]\} &=& \vec{D}[N_1\vec{\nabla}N_2-N_2\vec{\nabla}N_1]\,, \label{HH} \end{eqnarray} and are first-class (here, a metric is used to obtain the contravariant derivative $\vec \nabla$). The preservation of the constraints $\dot{C}[N_1]=\{C[N_1],H[N,\vec{M}]\}$ and $\dot{\vec{D}}[\vec{M}_1]=\{\vec{D}[\vec{M}_1],H[N,\vec{M}]\}$ then follows directly, and the gauge-invariant modes $\phi$ are those that satisfy $\{\phi,H[\epsilon_0,\vec{\epsilon}]\}=0$ (for all $\epsilon_0$ and $\vec{\epsilon}$) whenever the constraints hold. The relations (\ref{DD})--(\ref{HH}) have an interesting geometrical meaning \cite{Regained}: They realize commutators of deformations of spacelike hypersurfaces in space-time along vector fields $N\vec{n}+\vec{M}$ with unit normals $\vec{n}$ to the hypersurfaces. Although this algebra does not refer to the dynamics of the theory, it has a tight relation with it: Second-order field equations for the metric which are covariant under symmetries obeying the algebra (\ref{DD})--(\ref{HH}) must equal Einstein's equation (with an unrestricted cosmological constant) \cite{Regained,LagrangianRegained}. This result leaves only little room for quantum corrections. It is the canonical analog of the familiar statement that a local covariant action can correct the Einstein--Hilbert term only by higher-curvature contributions. When quantum corrections are inserted in the equations of gravity, in particular in cosmological perturbation equations, it is never clear whether the delicate consistency conditions summarized in the first-class nature of the constraint algebra remain intact. Especially background-independent frameworks cannot directly rely on standard covariance arguments because their notion of space-time, encoded in (\ref{DD})--(\ref{HH}), is supposed to emerge in some way from solutions to their equations. The consistency of the equations, however, must be ensured before they can be solved. If the approach used as well as the methods employed to derive semi-classical equations are covariant, the preserved symmetry implies consistent equations based on first-class constraints. This statement could, for instance, apply to effective equations derived from a path-integral quantization of gravity, provided one uses the correct integration measure to make the theory anomaly-free. (The latter condition is highly non-trivial.) In canonical quantum gravity, however, consistency must be shown explicitly because the different treatments of time and space derivatives in canonical quantizations eliminate manifest covariance. In our discussion so far we have assumed that the full field content of gravity is considered without fixing the space-time gauge. The issues we mention can formally be circumvented if one fixes the gauge before quantizing or before inserting quantum corrections in effective equations. In many cases, gauge fixing before quantization can indeed be assumed to be harmless, but the situation considered here is different. First, the constraints we are dealing with are more complicated functions than, say, the Gauss constraint of Yang--Mills theories. It is therefore more likely that the constraints receive significant quantum corrections. If the constraints are quantum corrected, the gauge transformations they generate are not of the classical form. Gauge fixing before quantization is then inconsistent, because one would fix the gauge according to transformations which subsequently will be modified. Secondly, in the present case the dynamics is part of the gauge system. A consistent theory must therefore quantize gauge transformations and the dynamics at the same time; one cannot fix one part (the gauge) in order to derive the second part (the dynamics) in an unrestricted way. Therefore, in canonical quantum gravity one must consider the full space-time gauge algebra without restrictions, or else one cannot be sure that the resulting theory is consistent. We should note that a less severe point of view, sometimes advocated in the literature, is to regard gauge fixing prior to quantization as part of the definition of a quantum theory which may turn out to be inequivalent to a non-gauge fixed quantum theory. In all cases where such a model can be checked to be self-consistent (as partially done in the hybrid cases mentioned in section\ \ref{hyb}), this attitude is legitimate. The resulting dynamics and physical predictions are, in general, quantitatively different from the theory quantized without fixing the gauge. Whenever available, however, the latter should be preferred because it implements the full system. Another method that avoids dealing with the full system (\ref{DD})--(\ref{HH}) is reduced phase-space quantization, often combined with a technique called deparametrization. (Examples for cosmological perturbations have been provided in \cite{PertObsI,PertObsII,BKdustI,BKdustII}.) For the reduced phase space, one solves the constraints classically and computes all observables invariant under gauge transformations. The resulting phase space is then to be quantized, which in general can be very complicated but may be possible in reduced or perturbative models. To facilitate the construction of gauge-invariant observables in a relational form, deparametrization selects a suitable phase-space degree of freedom to play the role of time. Formally, such a procedure leads to consistent equations, but they correspond to a quantized dynamics for classical observables rather than a complete quantization of the original gauge system. Moreover, for meaningful results one should check that the choice of internal time does not affect predictions, a form of covariance problem which is rarely analyzed in this context (see \cite{ReducedKasner}). The preferred method in our view is to quantize the constraints without classical specifications of gauge or observables. Canonical quantum gravity then provides operator versions of the classical constraints $C[N]$ and $\vec{D}[\vec{M}]$. Consistency is formulated just as for the classical constraints, except that commutators are used instead of Poisson brackets to define a first-class system. Making sure that such an operator algebra is first class is even more difficult than showing this property for a classical Poisson-bracket algebra because the result is usually sensitive to the particular factor ordering or possible regularization schemes used to define the operators. (A closed algebra may even be in conflict with Hermitian constraint operators \cite{Komar,NonHerm}.) It is therefore useful to combine the derivation of a consistent algebra with the one of effective or semi-classical equations. Such an approach cannot prove that any particular theory of quantum gravity is consistent, due to the more-involved nature of operator algebras. But it can show what kind of geometrical or physical implications quantized constraints can have, including the potential for observational tests in cosmology. With such an effective viewpoint, one (i) takes the corrections suggested by operator definitions in some approach to quantum gravity, (ii) parametrizes them so as to allow for sufficient freedom to encompass the ambiguities and unknowns in quantum operators, (iii) inserts them in the classical constraints and (iv) computes their algebra under Poisson brackets. In most cases, the algebra will no longer be first class, but careful choices of the functions and parameters used to specify the modifications may respect this important feature. If no choice of parameters gives rise to a first-class algebra, the quantum effect described by them is likely to be inconsistent. If there are consistent choices (which, in general, are not unique; see section\ \ref{fref}), it is possible for the quantum effect studied to be part of an anomaly-free theory of quantum gravity, and one obtains a consistent model whose equations can be analyzed for further implications. In particular, one can then derive a complete set of cosmological perturbation equations and observables, incorporating quantum corrections. (At this stage, consistency being assured, one may use gauge fixing of the quantum-corrected constrained system to simplify solving the equations of motion.) The prescription just sketched has been defined and evaluated in detail in \cite{ConstraintAlgebra}, with applications to a certain type of corrections suggested by loop quantum gravity. In these cases, the algebra of constraints remained first class, but it was deformed: Its structure functions in (\ref{HH}), not just the constraints themselves, were affected by quantum corrections. Two developments that have happened in the meantime motivate us to take a renewed look: First, canonical effective theories and especially the origin of higher time derivatives in them has been better understood. Secondly, applied to a different type of modifications in loop quantum gravity \cite{JR,ScalarHol}, the surprising (and in some eyes, shocking) possibility of signature change \cite{Action} at high density was found. As a consequence, the framework has been subjected to enhanced scrutiny and criticism. This article presents a unified setting, combining insights found in several papers since \cite{ConstraintAlgebra}, fleshing out some details that may so far have been mentioned only rather implicitly, and contrasting the approach with others that avoid dealing with the constraint algebra and instead use gauge fixing or deparametrization before the theory is quantized or modified. We end this introduction with a summary of our main results. This article includes several passages of review in order to make especially the conceptual part of our discussions self-contained. A clear-cut separation between review material and novel statements would be rather artificial, since both are interwoven and the former is systematically utilized to introduce the latter. For the reader's convenience, however, we state here the main results in a nutshell: \begin{enumerate} \item The role of quantum moments in the effective dynamics is clarified and contrasted with that of holonomy modifications. The main difference is that higher-order time derivatives arise only by the first type of corrections (section \ref{s:Corr}). Section \ref{s:Degrees} contains a brief summary of effective methods, with special emphasis on the anomaly problem and the generality of this type of effective constructions. The techniques of relevance for our considerations do not make use of further approximations, such as derivative expansions, which would restrict the availability of effective descriptions. \item By performing a detailed comparison of the most recent results on anomaly-free realizations of spherically symmetric models of loop quantum gravity as well as cosmological perturbations, we arrive at a simplified and more uniform picture of parametrizations of quantum corrections (section \ref{s:Examples}, especially sections \ref{323} and \ref{fref}). \item The definition and possibility of signature change, which is an important ingredient for understanding the early universe, is discussed at length in section \ref{s:Sig}. In particular, we show that signature change, if it occurs, is not a consequence of perturbative inhomogeneity but rather of holonomy modifications used crucially in the background dynamics of loop quantum cosmology. We also contrast signature change with the more common phenomenon of instability. \item Details are provided for systematic extensions of perturbation schemes to higher orders, paying due attention to degrees of freedom and symplectic structures (section \ref{s:Pert}). The role of gauge invariance and different applications of gauge fixing or deparametrization are discussed across various approaches to cosmological perturbations (section \ref{s:Add}). \item When computing the inflationary spectra, the choice of vacuum is simply a choice of initial conditions consistent with the canonical effective equations. It is not a ``re-quantization'' of the theory, as sometimes hinted at in the literature. Our discussion of this question clarifies this issue and has the additional advantage of generalizing the specification of vacuum initial conditions to non-standard space-time structures as encountered in models of loop quantum gravity (section \ref{s:InflVac}). \end{enumerate} | We have presented a detailed discussion of effective methods for models of canonical quantum gravity in relation to the anomaly problem. We have mainly discussed the overall coherence and consistency of the framework, but also provided new details and insights: (i) We have compared different anomaly-free models and showed their agreements, (ii) have considered higher-order perturbations in this framework, and (iii) have described an effective way of specifying an initial inflaton state. Our results pave the way to several new developments, all of which would require more time and space than is available for this article. However, while important technical details still have to be completed in these directions, the conceptual discussions presented here have removed what had appeared to be difficult hurdles. In particular, our description of a framework for anomaly-free higher-order perturbations will be important for attempts to derive non-Gaussianity from models of loop quantum cosmology. Specific calculations would require long manipulations of Poisson brackets of parameterized constraints with, say, holonomy modifications. If anomaly-free constraints existed to higher orders, their derivation would follow the scheme provided here. If such constraints did not exist, the same scheme would allow one to test whether new kinds of corrections could provide anomaly-freedom. Our characterization of an effective description of initial inflaton states should be easier to implement. Here, the main new direction would be an exploration of effects from non-vacuum states. More generally, the construction of an anomaly-free perturbation dynamics will permit the extraction of the full set inflationary observables in loop quantum cosmology, thus completing extant results based on a partial or preliminary implementation of quantum corrections. As emphasized, although a general framework to derive effective theories of canonical quantum gravity is available, it remains incompletely realized in present concrete calculations. Especially higher spatial and higher time derivatives must still be included in the existing models. No alternative procedure has yet yielded such terms for models of loop quantum gravity, and therefore it is unclear if they can be consistent at all. (If they cannot, there would be a significant mismatch between loop quantum gravity and general expectations from effective field theory, as realized in perturbative quantum gravity \cite{EffectiveGR,BurgessLivRev}.) Moreover, several features we described are lacking in such alternative approaches. We propose the effective approach sketched here as a candidate complete method to address the anomaly problem and to extract physical predictions from canonical quantum gravity. | 14 | 4 | 1404.1018 |
1404 | 1404.1532_arXiv.txt | We report the results of a deep {\it XMM-Newton} observation of the radio-faint $\gamma$-ray pulsar J1741$-$2054 and its nebula together with the analysis of five years of {\it Fermi} Large Area Telescope (LAT) data. The X-ray spectrum of the pulsar is consistent with an absorbed power law plus a blackbody, originating at least partly from the neutron star cooling. The nebular emission is consistent with that of a synchrotron pulsar wind nebula, with hints of spatial spectral variation. We extended the available {\it Fermi} LAT ephemeris and folded the $\gamma$-ray and X-ray data. We detected X-ray pulsations from the neutron star: both the thermal and non-thermal components are $\sim$35-40\% pulsed, with phase-aligned maxima. A sinusoid fits the thermal folded profile well. A 10-bin phase-resolved analysis of the X-ray emission shows softening of the non-thermal spectrum during the on-pulse phases. The radio, X-ray and $\gamma$-ray light curves are single-peaked, not phase-aligned, with the X-ray peak trailing the $\gamma$-ray peak by over half a rotation. Spectral considerations suggest that the most probable pulsar distance is in the 0.3-1.0 kpc range, in agreement with the radio dispersion measure. | \label{intro} The launch of the {\it Fermi $\gamma$-ray Space Telescope} offered the first opportunity to study a sizeable population of $\gamma$-ray pulsars. The {\it Fermi} Large Area Telescope \citep[LAT,][]{atw09} has discovered pulsed $\gamma$-ray signals from more than 150 objects \citep{abd13}, revolutionizing our view of them and giving birth to new high-energy pulsar sub-families, such as millisecond \citep[see e.g.][]{abd09a,ran11,kei11,esp13} and radio-quiet $\gamma$-ray pulsars \citep[see e.g.][]{abd09,saz10,ple13}, as numerous as the classic family of young, radio-loud pulsars \citep{car13}. The wealth of detections confirms the importance of the $\gamma$-ray channel in the overall energy budget of rotation-powered pulsars and paves the way to understanding the three-dimensional structure and electrodynamics of neutron star magnetospheres. Indeed, radio and $\gamma$-ray light curves contain a great deal of useful information about pulsar emission processes \citep[see e.g.][]{wat11,pie12,pie14}, confirming that models with emission originating at high altitudes in the magnetosphere \citep[e.g. outer and slot-gap,][]{che86,har04} are favored over models with near-surface emission \citep[e.g. polar cap,][]{har13}. Fitting $\gamma$-ray and radio light curves simultaneously is a promising way to constrain pulsar geometric parameters \citep[e.g.,][]{pie14}. Using the information in the (magnetospheric) non-thermal pulsar X-ray light curves could further improve the approach, adding another piece to the pulsar emission puzzle. This approach could localize the emitting region(s) responsible for the non-thermal pulsed X-ray emission with respect to the high altitude gamma-ray emitting region.\\ Few X-ray light curves have been exploited for modeling magnetospheric emission, compared to $\gamma$-ray profiles. This is largely due to the lack of high-quality X-ray light curves, primarily due to the occasional and non-targeted observational efforts. At this time, $\sim$60 out of 77 young pulsars in the second {\it Fermi} LAT pulsar catalog \citep[2PC,][]{abd13} have been detected in X-rays \citep{bec09,mar11,mar12,abd13}, but X-ray pulsations have been detected from fewer than half of them. Only nine {\it Fermi} pulsars have both the high X-ray fluxes and the long dedicated X-ray observations needed to disentangle the thermal and the non-thermal pulsations. Only five of these (Crab, Vela, Geminga, PSR J0659+1414 and PSR J1057$-$5226) have been characterized by a multi-bin phase-resolved X-ray spectral analysis \citep{del05,man07,wei11}. Of these only the Crab \citep[and possibly Geminga,][]{jac05} yielded a non-thermal light curve with a photon index varying with phase, a behavior that is still puzzling \citep{har08,tan08,hir08}. With the notable exception of the Crab among young pulsars, the multiwavelength behavior of isolated neutron stars is complex, with radio, optical, X-ray and $\gamma$-ray light curves usually unaligned, pointing to different emitting regions in the pulsar magnetosphere. The rich phenomenology, in particular including the X-ray information, has not yet been fully exploited for modeling the radiation processes of pulsars, leaving a number of questions unsolved. Here we report the results of a deep {\it XMM-Newton} observation intended for phase-resolved X-ray spectral analysis of the {\it Fermi} pulsar J1741$-$2054 (hereafter, J1741). The middle-aged J1741 ($\tau_c$ = 390 kyr) was discovered in a blind pulsation search of a {\it Fermi} LAT point source \citep{abd09}. For a moment of inertia $I = 10^{45}$ g cm$^2$, its period $P$ = 413 ms and period derivative $\dot{P}$ = $1.7\times10^{-14}$ s s$^{-1}$ give a spin-down energy loss rate $\dot{E}$ = 9$\times10^{33}$ erg s$^{-1}$, clearly on the low side of the $\gamma$-ray pulsar distribution \citep[see Figure 1 in ][]{abd13}. The pulsar was then detected in archival Parkes radio observations, showing a remarkably low dispersion measure, DM = 4.7 pc cm$^{-3}$ \citep{cam09}. The Galactic electron density model of \citet{cor02} yields a distance of $\sim$0.38 kpc, making J1741 one of the closest $\gamma$-ray pulsars known. At this distance, the low observed 1400 MHz radio flux density, S$_{1.4}$ $\sim$ 160 $\mu$Jy, makes it the faintest radio pulsar known. At the position obtained from LAT timing analysis \citep{ray11}, \citet{abd09}, \citet{rom10} and \citet{mar11} found the X-ray counterpart using both {\it Swift} and {\it Chandra} data. The {\it Chandra} observation also revealed diffuse, faint X-ray emission due to a pulsar wind nebula (PWN) trail extending some 2$'$ at position angle P.A. = 45$^{\circ}$ $\pm$ 5$^{\circ}$ (north through east). This extended structure was also associated with a 20$''$ long H$\alpha$ bow shock. Accurate bow shock modelling by \citet{rom10} suggests a pulsar velocity v$_p$ $\sim$ 150 km s$^{-1}$ directed 15$^{\circ}$ $\pm$ 10$^{\circ}$ out of the plane of the sky. | By analyzing the new {\it XMM-Newton} observation of PSR J1741$-$2054, we fully characterized the high-energy emission of this nearby middle-aged radio-faint pulsar.\\ Its nebular emission is typical of pulsar wind nebulae, both for its non-thermal spectrum, and for the flux decrease with distance along the tail emanating from the pulsar. A hint of nebular spectral variation with distance from the pulsar is detected. The shattered shape of the nebula is peculiar and a deeper analysis of the new {\it Chandra} data is needed to better understand this unusual behavior. These observations could also provide a measurement of the pulsar proper motion. Such an analysis is beyond the scope of this work. Modeling the X-ray spectrum of the pulsar requires a composite model, summing thermal and non-thermal components. Both components are $\sim$35-40\% pulsed, with single-peaked light curves and maxima phase-aligned to within $0.1$ in phase. While the thermal light curve is compatible with a sinusoid, the non-thermal profile has a sharper peak. The best-fitting thermal spectrum yields a temperature ($\sim7\times10^5$ K), compatible with the theoretical expectations for the cooling of a 390 kyr-old pulsar \citep{pon09}. We note that the best fit temperature of the {\tt nsa} model ($\sim3\times10^5$ K) is below the theoretical expectations, further disfavoring such a model. A pulsed component from thermal cooling has already been noted for several pulsars \citep{car04,del05,man07}, and that of PSR J1057$-$5226 has a similar pulsed component. Such pulsations can be ascribed to a dependence of the observed emitting area on the line of sight. \citet{hal93} describe a magnetospheric $``$blanket$"$ caused by cyclotron resonance scattering off the plasma in the magnetosphere that could screen the thermally emitting surface during specific phase intervals, depending on the magnetic field configuration and viewing geometry. Anisotropic heat transfer from the pulsar interior can also explain flux variation across the neutron star surface \citep{gre83}. If the thermal component is due to the cooling of the entire neutron star surface (12 km radius), from the thermal normalization in the best fit spectrum (that depends only on the pulsar distance and the emitting radius) we can derive the pulsar distance to be $\sim$850 pc, with 3$\sigma$ limits of $\sim$650 and $\sim$1100 pc \citep[e.g.,][]{hal07}. On the other hand, thermal emission from polar caps heated due to downstreaming of $e^{\pm}$ \citep{har02}, as seen in the case of PSR J0007+7303 in the CTA 1 supernova remnant, is expected to be generated from much smaller regions ($<$100 m, based on a simple $``$centered$"$ dipole magnetic field geometry, \citet{del05,mar13}), which would imply an unrealistic pulsar distance smaller than 10 pc. Distances lower than a few hundred parsecs are greatly disfavored due to the non-negligible value of the column density (about one fourth of the total Galactic column density in this direction). We note that a distance of 850 pc would result in a 110\% $\gamma$-ray efficiency, defined as $L_\gamma/\dot E$, with $L_\gamma$ the luminosity above 100 MeV. The distance range cited above implies an efficiency range of 60 to 180\%. Similarly, the X-ray efficiency would be $\sim$1.2\% (0.7 to 2.1\%). Such an unrealistic $\gamma$-ray efficiency could be explained by a beaming factor f$_{\Omega}$ less than 1 (as defined in 2PC), or by a moment of inertia larger than 10$^{45}$ g cm$^2$. A 10-bin phase-resolved X-ray spectral analysis reveals variations in the X-ray photon index, in addition to the phase-varying normalizations of the two spectral components, with a softer spectrum during the on-pulse phases (Figure \ref{fig5}). It is difficult to compare the behavior of J1741 with that of other {\it Fermi} pulsars, since the Crab is the only one for which such variation was detected, and the Crab's non-thermal spectrum becomes softer in the primary-pulse maximum and harder during the bridge between the two maxima \citep{wei11}. Although some models have been developed to explain the Crab's optical-to-$\gamma$-ray behavior \citep[see e.g.][]{har08,tan08,hir08}, the physics behind the pulsar's X-ray photon index variations is still unclear. Moreover, unlike the Crab, the $\gamma$-ray, X-ray and radio peaks of PSR J1741$-$2054 are not aligned, pointing to a clear difference in the geometry and/or altitude above the neutron star surface of the different emitting regions. Indeed, the $\gamma$-ray and X-ray peaks are phase-offset by roughly a half rotation, as is also seen in PSRs J0007+7303, J1057$-$5226 and J0659+1414 \citep{del05,car10}. Although these differences are expected from different models for the radio and $\gamma$-ray bands \citep{abd13}, no model is able to account for the offset between $\gamma$-ray and X-ray peaks. The alignment between the thermal and non-thermal X-rays (seen also in other pulsars, e.g. J0659+1414 and J1057$-$5226), as well as the phase lag with the $\gamma$-ray emission coming from the outer magnetosphere, can suggest that the non-thermal emission is generated in a region near the pulsar poles (e.g. in a polar cap emission model). Also, the low X-ray luminosity of radio-quiet pulsars in the X-ray band \citep{mar11} suggests that the radio and X-ray emission regions may be in close proximity. The origin of the phase lag between the radio and X-ray light curves is unclear. A comprehensive study of pulsar high-energy light curves and phase-resolved spectra will be crucial to understanding the X-ray emission mechanisms and geometry. \begin{figure*} \centering \includegraphics[angle=0,width=15cm]{fig1.pdf} \protect\caption{{\footnotesize Combined exposure-corrected 0.3-6 keV FOV images of the three {\it XMM-Newton} cameras. We applied a Gaussian filter with a kernel radius of 3$''$. Cyan symbols mark the sources detected with $>6\sigma$ confidence and with more than 225 detected counts that we analyzed to constrain the value of the Galactic absorption column. The squares mark the AGN-like sources, diamonds other field sources and the cross marks the pulsar. {\it Upper-left panel}: Expanded image of the pulsar and its tail. Green annuli mark the regions we used for the PSF analysis (with the exception of the nebular region); from white regions we extracted photons to build the pulsar and nebular spectra; red regions mark the pointlike sources we excluded from analysis.}} \end{figure*} \begin{figure*} \centering \includegraphics[angle=0,width=15cm]{fig2.pdf} \protect\caption{{\footnotesize Best-fit phase-resolved X-ray spectral parameters of PSR J1741$-$2054, plotted as a function of the pulsar phase, defined as in Figure 3. As discussed in Section~\ref{spres}, the power law (panel 3 from the top) and blackbody (panel 4) normalizations and the photon index (panel 2) evolve throughout the pulsar phase.}} \end{figure*} \begin{figure*} \centering \includegraphics[angle=0,width=15cm]{fig3.pdf} \protect\caption{{\footnotesize {\it Top Panel}: EPIC/pn folded light curves in different energy ranges. X-ray photon phases were computed according to the {\it Fermi} LAT ephemeris overlapping with the {\it XMM-Newton} dataset, with selection as in Section 6. The red curve contains photons in the 0.3-0.7 keV energy range and the green one in the 0.7-10 keV range. The curves have been renormalized by dividing each bin by N$_{counts}$/N$_{bins}$, where N$_{counts}$ is the total number of events in the energy range and N$_{bins}$ the number of bins (1$\sigma$ errors are shown). {\it Bottom Panel}: Phased radio, X-ray, and $\gamma$-ray light curves of PSR J1741$-$2054. The 300 MHz radio light curve, shown in blue, comes from the Green Bank Observatory \citep{cam09}. The 100-bin $\gamma$-ray curve, shown in black, contains all the 5-year {\it Fermi} LAT weighted counts with energies $>$100 MeV. The 20-bin X-ray curve is shown in cyan. All the curves have been renormalized to have the highest bin value equal to 1 (1$\sigma$ errors are shown).}} \end{figure*} \begin{figure*} \centering \includegraphics[angle=0,width=15cm]{fig4.pdf} \protect\caption{{\footnotesize Confidence contours for the 10-bin X-ray phase-resolved analysis of PSR J1741$-$2054, showing the pulsar photon index and the power law normalization. Black contours are at the 1$\sigma$ confidence level, red at 90\% and green at 3$\sigma$. Blue lines follow the pulsar phase.}} \end{figure*} | 14 | 4 | 1404.1532 |
1404 | 1404.4121_arXiv.txt | We investigate the mechanism to reproduce notable spectral features at the ignition phase of nova explosion observed for a super-Eddington X-ray transient source MAXI J0158$-$744 in the Small Magellanic Cloud. These are a strong Ne IX emission line at 0.92 keV with a large equivalent width of $0.32^{+0.21}_{-0.11}$ keV and the absence of Ne X line at 1.02 keV. In this paper, we calculate the radiative transfer using a Monte Carlo code, taking into account the line blanketing effect due to transitions of N, O, Ne, Mg and Al ions in an accelerating wind emanating from a white dwarf with a structure based on a spherically symmetric stationary model. We found that the strong Ne IX line can be reproduced if the mass fraction of Ne is enhanced to $10^{-3}$ or more and that of O is reduced to $\sim5\times10^{-9}$ or less and that the absence of other lines including Ne X ions at higher energies can be also reproduced by the line blanketing effect. This enhancement of the Ne mass fraction indicates that the ejecta are enriched by Ne dredged up from the surface of the white dwarf composed of O, Ne, and Mg rather than C and O, as already pointed out in the previous work. We argue that the CNO cycle driving this nova explosion converted most of C and O into N and thus reduced the O mass fraction. | MAXI (Monitor of All-sky X-ray Image; \citealt{2009PASJ...61..999M}) discovered a luminous soft X-ray transient, MAXI J0158$-$744, in the direction of the east edge of the Small Magellanic Cloud (SMC) on November 11, 2011 \citep{2011ATel.3756....1K}. The spectra of the initial X-ray flash are reproduced by blackbody models with the temperature of $\sim300$ eV and the photospheric radius of $\sim2,000$ km \citep{2013ApJ...779..118M}. The Swift follow-up observations, which started from 0.44 days after the MAXI discovery, revealed that the X-ray spectra, which are fitted by blackbody spectra at temperature $\sim0.1$ keV, were similar to those of novae in the super-soft source (SSS) phase \citep{2012ApJ...761...99L}. The optical spectra taken by SAAO, ESO, and SMARTS telescopes indicated the characteristics of a Be star at the distance of the SMC \citep{2012ApJ...761...99L,2013ApJ...779..118M}. Thus, MAXI J0158$-$744 was recognized as a nova accompanied by a Be star. The X-ray flash detected by MAXI GSC (Gas Slit Camera; \citealt{2011PASJ...63S.623M,2011PASJ...63S.635S}) and SSC (Solid-state Slit Camera; \citealt{2010PASJ...62.1371T,2011PASJ...63..397T}) lasted for $\sim1300$ s at least. The maximum luminosity was $\sim100$ times brighter than the Eddington limit of a solar mass object. MAXI J0158$-$744 behaved differently from most novae known up to the present. While standard novae emit optical photons for at least several days before emitting soft X-ray photons during the SSS phase, the outburst of MAXI J0158$-$744 was first detected in the X-ray band, not in optical. \citet{2013ApJ...779..118M} reported that the unabsorbed X-ray luminosity was about $3\times10^{39}$ erg s$^{-1}$ at $t=8$ s ($=t_{8}$) in the energy range of 0.7$-$7.0 keV, assuming the distance of 60 kpc (Hereafter, the origin of time is referred to the MAXI trigger time.). The X-ray luminosity reached the peak of $L_{\rm p}\sim2 \times 10^{40}$ erg s$^{-1}$ at $t=220$ s ($=t_{220}$), and then decreased to $L_{1296}\sim7 \times 10^{39}$ erg s$^{-1}$ at $t=1296$ s ($=t_{1296}$). Since these luminosities exceed the Eddington luminosity for a solar mass object by a factor of more than 10, a strong wind is expected to blow from the surface of the white dwarf (WD). This source is also characterized by a lack of the early optical phase and the very fast decay of the initial X-ray flash. The scan periods of the MAXI detectors constrain the duration $\Delta t$ of the event to be less than $1.10\times10^{4}$ s. If the energy emitted in the X-ray band is supplied by hydrogen burning, we can estimate the mass of the exhausted hydrogen fuel as $M_{\rm H}=L_{\rm p}\Delta t/0.007c^{2}\sim2\times10^{-8}M_{\odot}$, which is much smaller than those in usual novae ($\gtrsim10^{-7}M_{\odot}$; \citealt{1982ApJ...257..767F}). Because of this tiny mass, the wind emanating from the WD becomes transparent soon after the thermonuclear runaway (TNR) outbreak. Especially, the most remarkable feature is the X-ray spectrum taken by MAXI/SSC at $t=t_{1296}$, which exhibits a strong He-like Ne (Ne IX) emission line at the energy of 0.92 keV and no prominent line of H-like Ne (Ne X) at 1.02 keV \citep{2013ApJ...779..118M}. We adopt the results of the analysis by \citet{2013ApJ...779..118M} using blackbody fits, defining $R_{1296}=2,290$ km and $k_{\rm B}T_{1296}=0.33$ keV (where $k_{\rm B}$ is the Boltzmann constant) as fiducial values of the photospheric radius and temperature at $t=t_{1296}$. Though \citet{2012ApJ...761...99L} tried to explain the super-Eddington luminosity by a shock induced model, \citet{2013ApJ...779..118M} pointed out the difficulty in such a shock-induced model. The latter authors instead argued that the observed phenomenon corresponds to the fireball phase at the ignition of a nova explosion and that the luminosity is possibly explained by the convection on the surface of the WD, releasing a large amount of energy produced by the TNR during the first 100 s. They inferred that the He-like Ne line can be explained by the optically thin region around the photosphere. In their work, optically thin emission component with a temperature below 0.3 keV (Mekal component in \citealt{2013ApJ...779..118M}) and exceptionally large Ne abundance of 10 solar or more was necessary to explain the strong He-like Ne line and the lack of the H-like Ne line. In this case, a large emission measure ($\sim10^{63}$ cm$^{-3}$) is necessary and might be incompatible with the assumption of optically thin emission. In this paper, we investigate the possibility that photons scattered by N, O, Ne, Mg, and Al ions in a supersonic wind would exclusively form a strong K$\alpha$ line of He-like Ne and H-like N with P-Cygni profiles in the spectrum taken at $t=t_{1296}$ and eliminate other lines by line blanketing effects. The line blanketing occurs when the Doppler broadened line profile overlaps with that of another line at a higher energy or shorter wavelength, and weakens the latter line and strengthen the former line. A supersonic wind is expected to enhance this effect owing to large Doppler shifts of the line energies. We use a Monte Carlo method to calculate the photon transfer and try to reproduce the spectral features at $t=t_{1296}$. In Section \ref{sec:formulation}, we present our simplified stationary wind model including the velocity and density profiles and describe the procedure to determine the mass loss rate from observed quantities. The radiative processes involved in our Monte Carlo method are presented in Section \ref{sec:MCcode}. In Section \ref{sec:results}, we show the results and compare them with observations. We conclude this paper in Section \ref{sec:conclusions}. | \label{sec:conclusions} We have examined whether resonance scattering off N, O, Ne, Mg, and Al ions in an accelerating wind reproduces a strong Ne IX K$\alpha$ emission line and lack of the Ne X K$\alpha$ line in the observed X-ray spectrum of MAXI J0158$-$744, by performing Monte Carlo calculations. We have constructed a simplified stationary model for a given luminosity with the ion distribution calculated using the XSTAR subroutines. The mass loss rate is determined to reproduce the observed photospheric radius. As a result of a series of the calculations presented here, the observed spectral features can be reproduced by emission from the expanding wind irradiated by a super-Eddington soft X-ray source. It is found that the Ne mass fraction needs to be enhanced by a factor of several tens of times compared with the SMC abundance ($X_{\rm Ne}\sim1\times10^{-2}$) to reproduce the observed Ne line at 0.92 keV. This amount of Ne could be supplied by matter dredged up from the surface of a WD composed of ONe. It should be noted that the ionization states estimated by ignoring the advection in the wind might overestimates the ionization degree and thus the necessary mass fraction of Ne. % We also find that the inclusion of Ne X ions at energies higher than 0.92 keV does not induce the Ne X K$\alpha$ line due to the line blanketing. In the same way, inclusion of heavier elements like Mg and Al supply more photons to the Ne IX K$\alpha$ line and strengthen it. Meanwhile, even a small amount of O weakens the Ne emission line for the same reason. It is necessary to decrease the mass fraction of O under $5\times10^{-9}$ for removing the influence on the Ne line completely. We have shown that the CNO cycle beneath the photosphere can reduce the amount of O. To reduce its mass fraction under $5\times10^{-9}$, the duration of the CNO cycle is constrained to be shorter than 4430 s. The inferred hydrogen mass burned out by the CNO cycle is $\approx10^{-8}M_{\odot}$. This is far smaller than the minimum accretion mass that enables ignition in hydrostatic equilibrium, $5\times10^{-7}M_{\odot}$ \citep{1982ApJ...257..767F}. Taking this into account, hydrogen should be accreted dynamically onto the surface of the WD and the subsequently generated shock ignites hydrogen leading to TNR. Such a dynamical behavior may be responsible for the dredge up of Ne ions, which is needed to explain the observed Ne line. The mass of such a WD must be very close to the Chandrasekhar limit, which was also indicated from the small photospheric radii deduced from observations \citep{2013ApJ...779..118M}. | 14 | 4 | 1404.4121 |
1404 | 1404.1642_arXiv.txt | We study the environments of low redshift ($z < 0.5$) quasars based on a large and homogeneous dataset from the Stripe 82 region of the Sloan Digital Sky Survey (SDSS). We have compared the $< 1$ Mpc scale environments of 302 quasars that were resolved in our recent study to those of 288 inactive galaxies with closely matched redshifts. Crucially, the luminosities of the inactive galaxies and the quasar host galaxies are also closely matched, unlike in most previous studies. The environmental overdensities were studied by measuring the number density of galaxies within a projected distance of 200 kpc to 1 Mpc. The galaxy number density of the quasar environments is comparable to that of the inactive galaxies with similar luminosities, both classes of objects showing significant excess compared to the background galaxy density for distances $<$ 400 kpc. There is no significant dependence of the galaxy number density on redshift, quasar or host galaxy luminosity, black hole mass or radio loudness. This suggests that the fueling and triggering of the nuclear activity is only weakly dependent on the local environment of quasars, and the quasar phase may be a short-lived common phase in the life cycle of all massive galaxies. Key words: galaxies: active - galaxies:nuclei - quasars: general | The last decades have seen the emergence of the general consensus that most, if not all, massive galaxies host a supermassive black hole in their center \citep[e.g.][]{richstone98}. Observations of early-type galaxies have shown a tight relation between the mass of the central black holes and the properties of the spheroids hosting them \citep[see e.g.][for a review]{ferrarese06} which suggests that the formation and evolution of the galaxies and their nuclear activity are linked. As quasars are fueled by accretion onto the supermassive black hole \citep[e.g.][]{yu02}, understanding the mechanism that triggers their activity plays a fundamental role in our understanding of the processes that have built the galaxies and their nuclei. In spite of half a century of studies aimed to understand the quasar activity, the mechanism that activates and fuels the nuclei of galaxies is still a matter of debate. The leading processes thought to be responsible for transforming a dormant massive black hole into a luminous quasar are dissipative tidal interactions and galaxy mergers \citep[e.g.][and references therein]{dimatteo05,callegari11}. Galaxy formation is known to be heavily influenced by the environment, with galaxies in clusters tending to be elliptical and deprived of most of their gaseous content \citep[e.g.][]{silk93,kormendy09}, and also commonly showing signs of close interactions and mergers \citep[e.g.][]{bennert08,mcintosh08}. To better understand how quasars are formed, it is therefore important to study the relation between this environment and the nuclear activity. The environments of quasars have been studied in the past on widely different scales ranging all the way from host galaxies to Mpc scales. Very early studies such as \citet{stockton78} and \citet{yee84} have shown that typical quasar environments have galaxy densities comparable to galaxy groups or poor clusters. At Mpc scales, comparing the environments of quasars to those of galaxies has given conflicting results. Early studies on Mpc scales suggest that quasars are more strongly clustered than galaxies \citep[e.g.][]{shanks88,chu88}, while later studies based on surveys such as the Two Degree Field (2dF) and the Sloan Digital Sky Survey (SDSS) have found the environmental galaxy densities of quasars and galaxies to be comparable to each other \citep[e.g.][]{smith00,wake04}. At smaller scales, the studies have also shown differing results. \citet{ellingson91} studied a sample of 32 radio loud quasars (RLQs) and 33 radio quiet quasars (RQQs) at $0.3 < z < 0.6$ and found that the environments around RLQs are significantly denser than those around RQQs, which they find to have environmental density values similar to those found for average non-active galaxies in previous studies. More recent studies of small samples of low redshift quasars such as \citet{fisher96} and \citet{mclure01} on the other hand find no difference between the environments of RLQs and RQQs. Both studies used data taken with the HST, with \citet{fisher96} studying a sample of 20 quasars at $z \leq 0.3$, and \citet{mclure01}) using a sample of 44 quasars at $z ~0.2$. The environments of both RLQs and RQQs were found to have densities larger than those of non-active galaxies, with values similar to those found for the RLQ sample of \citet{fisher96}. Similar results were reported by \citet{wold00} and \citet{wold01} who used images from Nordic Optical Telescope (NOT) and HST to study 21 RLQs and 21 RQQs at $0.5 < z < 0.8$, concluding that no evolution of the environmental density with redshift was found. A more recent study by \citet{almeida13} compared the environments of 19 radio galaxies at redshifts $0.2 < z < 0.7$ to those of 20 RQQs at $0.3 < z < 0.41$ and found that radio galaxies appear to reside in denser environmets than quasars. The early studies of quasar environments had to deal with small sample sizes, but surveys such as the Two Degree Field (2dF) and Sloan Digital Sky Survey (SDSS) have allowed studies with much larger quasar and galaxy samples. \citet{croom04}, for example, used data from the 2dF QSO Redshift Survey (2QZ) to study the clustering of of $~20000$ quasars at $z < 3$. The study of the actual environments of the quasars and a comparison sample of galaxies was done at redshifts $z < 0.3$ for a subsample of $~200$ quasars, and \citet{croom04} found the environments of the quasars to be statistically identical to those of galaxies. \citet{coldwell06} used the third data release (DR3) of SDSS to study the environments of $~2000$ quasars at redshifts of $z < 0.2$, using a comparison sample of 2300 galaxies. The quasar and galaxy samples were selected to have similar redshift distributions, but no matching based on the luminosity of the galaxies was done. The study found no difference between the environmental density of the two samples, noting that both quasars and galaxies tend to reside in regions more dense than field galaxies, but less dense than cluster environments. However, some other studies using the SDSS archives have found contradicting results. \citet{serber2006} also used data from SDSS DR3 to study environments of $~2000$ quasars at $z < 0.4$ and found again that quasars are located in regions of local overdensity higher than that of the background and that the density enhancement is strongest within 100 kpc from the quasar. They also found that the overdensity around the quasars is larger than that around typical L* galaxies, and that the high luminosity quasars have denser small-scale environments than low luminosity quasars. Another study by \citet{strand2008} used SDSS DR5 to study a sample of ~4000 quasars at $z < 0.6$, found environmental densities similar to those of \citet{serber2006}, though a comparison with non-active galaxies could not be performed due to the lack of a control sample. A more recent study by \citet{zhang13} used data from Stripe 82 region of the SDSS to study environments of 2300 quasars at $0.6 < z < 1.2$. They found that quasars exhibit an overdensity of galaxies with respect to the background of field galaxies, and that the clustering amplitude increases with the redshift. However, in the absence of a control sample, it is unclear whether the density of galaxies around quasars actually differs from the density around non-active galaxies at these redshifts. In this study we have used more recent data from the seventh SDSS data release (DR7) \citep{abazajian09} to study the environments of low redshift quasars. In order to perform the environment study with deeper capability, we have used a stripe of sky along the Celestial Equator in the Southern Galactic Cap known as Stripe 82 \citep{annis2011}. This region was imaged multiple times during the period of 2004-2007, and the final co-added images reach up to 2 magnitudes deeper than other SDSS data. This allowed us to use a fainter magnitude limit for the galaxies to be included than the previous studies. This work is part of a series of papers aimed at investigating the properties of low redshift quasars from a large and homogeneous dataset. In paper I \citep{falomo2014} the results on the properties of quasar host and the relationship with BH mass were reported. In paper III (Bettoni et al. in prep) we aim to study the morphology, colours and peculiarities of quasar hosts, while paper IV will study the colours of galaxies in the environments of quasars. The work done in paper I allowed us to construct a control sample of galaxies which is closely matched with the quasar sample with respect to the properties of the host galaxies, an important aspect which was neglected by previous studies. The quasar and galaxy samples used in the study are described in Section 2. The dataset and the method of analysis are described in Section 3, and the results are given in Section 4. Section 5 contains comparison with previous works. Finally in Section 6 we report the main conclusion of this study. We adopt the concordance cosmology with H$_0$ = 70 km s$^{-1}$ Mpc$^{-1}$, $\Omega_m$ = 0.3 and $\Omega_\Lambda$ = 0.7. | We have used SDSS Stripe 82 data to study the $<1 Mpc$ environments of low redshift ($0.1<z<0.5$) quasars, and compared them with the environments of galaxies in the control sample of non active galaxies. The environments were studied by measuring the number density of galaxies within a projected distance of 200 kpc to 1 Mpc from the quasars, and then dividing this by the number density of galaxies in the background to calculate the overdensity of the region. The overdensities associated with the quasar environments were then compared with those of a sample of control galaxies well matched in both the redshift and the galaxy luminosity. We find the following results: \begin{enumerate} \item Quasars are on average found associated with small group of galaxies. The overdensities of galaxies are mainly observed in the closest ($<$ 200 kpc) region around the source and vanish at distance of 1 Mpc. \item No statistically significant difference is found between the overdensities around the quasars and the inactive galaxies at any separation. \item No dependence of the overdensity on redshift, quasar luminosity, the luminosity of the host galaxy, black hole mass or the radio luminosity was found. The result is the same for both our full and the matched samples. \end{enumerate} We compared our results to those of previous studies, in particular the ones by \citet{serber2006} who studied low redshift quasar environments ($z < 0.4$) and \citet{zhang13} who focused on redshift of $0.6 < z < 1.2$. We find that the galaxy overdensity is independent of either redshift or luminosity. For our sample of inactive galaxies we find a trend of higher overdensities with increasing luminosity. Contrary to \citet{serber2006}, we find the environments of quasars are similar to those of inactive galaxies. Additionally, the overdensities we find are lower for both the quasars and galaxies than those of \citet{serber2006}. The fact that we find no significant difference between the environments of quasars and non-active galaxies suggests that the link between the quasar activity and the environment of the quasars is less important than believed for fueling and trigerring the activity. This also points to smaller importance of (major) mergers than expected in triggering the quasar activity, and that secular evolution (e.g. disk instabilities) may play an important role. Similarity between the environments of quasars and non-active galaxies could thus indicate that the quasar phase is a common event in the life cycle of a massive galaxy and does not depend significantly on the local environment at least on the scales studied here. We also find no dependence of the radio loudness of the quasar on the overdensity, in agreement with studies by \citet{fisher96} and \citet{mclure01}. However, as our sample of RLQs is very small, and their radio luminosities quite low, we cannot make rule out the possibility connection between the environment and radio activity, such as the one found in \citet{almeida13}. A detailed study of morphology, peculiarities of quasar hosts and colours of this large sample of AGN could provide further clues to understanding the link between nuclear activity and the processes fueling and triggering it. This will be explored in a future paper of this series (paper III, Bettoni et al. in prep). Finally, the colours of the galaxies in the environments of the quasars in this sample will be explored in paper IV of the series (Karhunen et al. in prep). | 14 | 4 | 1404.1642 |
1404 | 1404.6595_arXiv.txt | The Yakutsk array group is developing a wide field-of-view Cherenkov telescope to be operated in coincidence with the surface detectors of the extensive air shower array. Currently, the engineering prototype of the reflecting telescope with the front-end electronics is designed, assembled, and tested to demonstrate the feasibility of the conceived instrument. The status and specifications of the prototype telescope are presented, as well as the modernization program of the already existing Cherenkov light detectors subset of the array measuring ultra-high energy cosmic rays. | Investigation of Cherenkov light induced by cosmic rays (CRs) cascading in the atmosphere began in the middle of the last century in the UK and the USSR. An exhaustive description of the early developments of the study of Cherenkov light is provided in \cite{Jlly}, while recent reviews of the entire area are presented, e.g., in \cite{Wtsn,Mrzn}. It is now well-known that the angular and temporal structure of the Cherenkov light emitted by an extensive air shower (EAS) can be used to infer the longitudinal development parameters of the shower; specifically, the lateral distribution of the light intensity measured is often used to estimate the energy and mass of the primary particle initiating EAS \cite{Ztspn, Fmn, Klmkv, Trvr}. The angular distribution of Cherenkov photons from EAS was calculated by V.I. Zatsepin~\cite{Ztspn} assuming it is determined primarily by that of electrons in the shower. Subsequently, Fomin and Khristansen proposed~\cite{Fmn} to use the pulse shape of the Cherenkov signal, namely, the pulse width, to indicate the shower's maximum position, $x_m$, in the atmosphere. Experimental measurements of the Cherenkov signal pulse shape were performed initially in Yakutsk and in Haverah Park \cite{Klmkv, Trvr}. The results were used to estimate $x_m$, and attempts were made to evaluate the cascade parameters of electrons at CR energies of approximately $10^{17}$ eV \cite{Trvr,Prsn}. A variety of detectors were used then; for example, the Tunka experiment operates an array of Cherenkov detectors near Lake Baikal \cite{Tunka2}. Our intention to develop a Cherenkov telescope functioning as a differential detector of EAS was motivated by the possibility to measure the depth of the cascade maximum and/or the shower age via the angular and temporal distributions of the Cherenkov signal \cite{NIMA}. Combining $x_m$ and the shower age with other characteristics measured with surface detectors of the EAS array, e.g. the energy and muon content, one is able to estimate the average mass composition of CRs. Experimental arguments in elucidating the origin of the knee and ankle in the CR spectrum will be significantly strengthened by the measurements of the angular and temporal distributions of the Cherenkov signal in the energy range above $10^{15}$ eV. Existing scenarios of CR acceleration in the sources differ in the expected model composition around the knee and in the transition region between galactic and extragalactic components \cite{Brzk}, so accurate estimation of the average mass of CR nuclei, in addition to the improved measurement of the sharpness of the knee and ankle, should allow us to discriminate some scenarios. The paper is structured as follows. The results of Cherenkov light modeling in EAS are described in Section 2. The design and performance of the wide field-of-view (FOV) telescope prototype is described in Section 3. The first results of EAS measurements with a prototype working in coincidence with the surface detectors are presented in Section 4. Our intentions for the modernization of the Cherenkov detectors of the Yakutsk array are discussed in Section 5. A summary is provided in Section 6, followed by three Appendices. | We designed and assembled the engineering prototype of the WFOV Cherenkov telescope to operate in cooperation with the surface detectors of the Yakutsk array. Field testing of the telescope demonstrated the practical applicability of the ideas proposed; the performance agrees with the main expectations from simulations. During the winter of 2013-2014, a number of EAS events were detected by the telescope in coincidence with the surface detectors subset of the Yakutsk array. The detection efficiency of the telescope was measured, as well as the effective radius of the telescope detecting area. The angular and temporal structural parameters of the air Cherenkov light signal from the particular EAS events were measured. A way of modernizing the Yakutsk array is proposed by adding a set of WFOV telescopes. In conjunction with the fast synchronization system of the array detectors, it will provide quite new possibilities in measuring the mass composition of cosmic rays in the energy range above $10^{15}$ eV. \begin{table}[t]\centering \caption{Estimation of the $x_m^{Cher}$ reconstruction accuracy, g/cm$^2$. $E_0=10^{15}$ eV.} \begin{tabular}{rrrr} \hline $R$, m & 600 & 400 & 200 \\ $\sigma_t= 3$ ns & 5.8 & 9.1 & 12.3 \\ $\sigma_t=10$ ns & 19.4 & 30.2 & 41.0 \\ $\sigma_R= 5$ m & 5.3 & 7.0 & 9.0 \\ $\sigma_R=50$ m & 53.0 & 70.0 & 89.9 \\ \hline \end{tabular} \label{Table:3}\end{table} | 14 | 4 | 1404.6595 |
1404 | 1404.4245_arXiv.txt | The ejecta discharged by impacting meteorites can redistribute a planetary ring's mass and angular momentum. This `ballistic transport' of ring properties instigates a linear instability that could generate the 100--1000-km undulations observed in Saturn's inner B-ring and in its C-ring. We present semi-analytic results demonstrating how the instability sustains steadily travelling nonlinear wavetrains. At low optical depths, the instability produces approximately sinusoidal waves of low amplitude, which we identify with those observed between radii 77,000 and 86,000 km in the C-ring. On the other hand, optical depths of 1 or more exhibit hysteresis, whereby the ring falls into multiple stable states: the homogeneous background equilibrium or large-amplitude wave states. Possibly the `flat zones' and `wave zones' between radii 93,000 and 98,000 km in the B-ring correspond to the stable homogeneous and wave states, respectively. In addition, we test the linear stability of the wavetrains and show that only a small subset are stable. In particular, stable solutions all possess wavelengths greater than the lengthscale of fastest linear growth. We supplement our calculations with a weakly nonlinear analysis that suggests the C-ring reproduces some of the dynamics of the complex Ginzburg--Landau equation. In the third paper in the series, these results will be tested and extended with numerical simulations. | The component particles of planetary rings suffer a continual bombardment of hypervelocity meteoroids, the impacts of which liberate a significant amount of material. Typically, impact ejecta reaccrete on to the ring but at a different radius from where they originated; they hence redistribute its mass and angular momentum. This `ballistic transport' of ring properties occurs on a characteristic lengthscale $l_\text{th}\sim 10-10^3$ km (the `throw length') and a timescale $t_e \sim 10^5-10^7$ yr (the `erosion time') (Durisen 1984, Ip 1984, Lissauer 1984). Other than influencing the large-scale evolution of Saturn's rings, ballistic transport instigates a linear instability that can spontaneously create structure on these scales (Durisen 1995). It has been argued that the 100-km waves in the inner B-ring and the 1000-km undulations in the C-ring are a result of the instability's nonlinear saturation (Durisen et al.~1992, hereafter D92, Durisen 1995, Charnoz et al.~2009, Colwell et al.~2009). This is the second paper in a series devoted to the dynamics of the ballistic transport instability (BTI) and its generation of axisymmetric structure. The first paper, Latter et al.~(2012) (hereafter Paper 1), outlined a convenient theoretical framework within which to attack the problem and rederived the BTI's linear theory. Here we aim to go further by tracking the BTI's nonlinear saturation. Ultimately, one is obliged to numerically simulate its evolution, and we present such calculations in the third paper of the series (Latter et al.~2013, submitted, hereafter Paper 3). In this work, however, we take a dynamical systems approach and establish a set of `a priori' results that can both guide and explain the simulations. First we demonstrate that ballistic transport supports families of steadily travelling nonlinear wavetrains. These solutions may be computed directly from the system's governing evolution equation. At low optical depths $\tau$, the wavetrains assume small amplitudes and possess approximately sinusoidal profiles.We identify them with the long 1000-km undulations in the C-ring between radii 77,000 and 86,000 km (see Fig.~13.17 in Colwell et al.~2009), but conclude that the 100-km plateaus at slightly larger radii are not generated by the BTI, at least not working in isolation. Meanwhile, when $\tau\gtrsim 1$ the system exhibits hysteresis: the homogeneous state is linearly stable, but there exist additional wave solutions of large amplitude. This raises the possibility that stable homogeneous states spatially adjoin stable wave states, with the interfaces possibly undergoing their own dynamics.This theoretical scenario compares well with observations of `flat' and `wave' zones in the inner B-ring between radii 93,000 and 98,000 km (see Fig.~13.13 in Colwell et al.~2009). For sufficiently small viscosities, hysteresis extends to very large optical depths. In fact, one can find nonlinear BTI-supported waves for $\tau > 2.5$, though it is unlikely such structures are relevant to ring observations. Subsequently, we determine the linear stability of these solutions and find that only a small subset are stable. As stable solutions possess wavelengths longer than that of the fastest growing linear mode, it is likely that the system undergoes a wavelength selection process, whereby power initially localised to the most unstable lengthscale seeks out the longer stable wavetrain solutions. Finally, we conduct a weakly nonlinear analysis of the long and slow dynamics of wavetrain modulations. It turns out that the wave amplitudes obey the complex Ginzburg--Landau equation, which suggests that the C-ring undulations share some of its non-trivial dynamics. The structure of the paper is as follows. In the following section, we summarise the relevant contents of Paper 1, such as the governing mathematical formalism, main parameters, and the BTI's linear stability analysis. In Section 3 we calculate the nonlinear wavetrain solutions, focussing on the two parameter regimes associated with the C-ring and inner B-ring. Section 4 outlines the linear stability of these structures, while Section 5 and the Appendix present a weakly nonlinear analysis of their long and slow modulations. We bring together these various results in the final Discussion section. | In this final section we summarise our results and apply them to the observational problems of Saturn's B- and C-rings. We also point towards future work. First, we have shown that the BTI can saturate via the formation of steadily travelling nonlinear wavetrains. Near marginal stability at low $\tau$, these solutions inhabit an interval of intermediate wavenumber $q$. For example, when the mean optical depth $\tau_0=0.175$, wavetrains exist with a $q$ between 1.75 and 3.59 (in units of $1/l_\text{th}$). The amplitudes of these waves are relatively small, with $\tau$ varying by $\sim 0.1$ between peaks and troughs (cf.\ Fig.~\ref{Cringprofs}). For the most part, these solutions are close to sinusoidal in appearance and possess phase speeds approximately equal to the linear BTI modes (Fig.~\ref{Cringvsq}). On the other hand, at large $\tau$ the system permits hysteresis: even if a homogeneous ring is linearly stable it can still support large-amplitude wavetrain solutions via the ballistic transport mechanism. Consequently, the ring will want to evolve to either the flat homogeneous state, or a stable wave state. The wavetrains do not resemble sinusoids, and the peak to trough variation is large, varying between 1 and 1.6 in $\tau$ (Fig.~\ref{BringProf}). Wavecrests in this marginal high-$\tau$ parameter regime propagate extremely slowly, at most with a phase speed $\sim 0.01 \,l_\text{th}/t_e$ or 1-100 mm yr$^{-1}$ (see Fig.~\ref{Bringvsq}). We tested the linear stability of these structures and found that for given parameters only a subset of the wavetrain solutions are stable. Generally, stable solutions possess the greatest amplitude and propagate the slowest. It is likely that the system will select one of these solutions if left to freely evolve. We also demonstrate that low-amplitude wavetrains undergo large-scale modulations which are governed by the complex Ginzburg-Landau equation. The amplitudes of our C-ring waves may then share in its interesting, sometimes disordered, dynamics. These results are compatible with observations of B- and C-ring structure (Porco et al.~2005, Colwell et al.~2009), as well as previous simulations of the inner B-ring (D92). Turning to the B-ring first, it is likely that the observed adjoining flat and wave zones between 93,000 and 98,000 km (Fig.~13.13 in Colwell et al.~2009) are products of the hysteresis exhibited by our model. The flat regions correspond to where the ring has fallen into the stable homogeneous state, and the wave regions to where it has jumped into the stable wave state. These zones are connected by fronts, which should exhibit additional dynamics that time-dependent simulations may probe. Perturbations that may have thrust B-ring regions out of the homogeneous state's basin of attraction might include the inner B-ring edge, the transition to extremely large $\tau$ at $r=99,000$ km, or the Janus/Epimetheus 2:1 inner Lindblad resonance. There are two problems that this scenario faces. First is the deepness of the troughs in the theoretical wave profiles. Typically the theoretical troughs possess a dynamical optical depth of $\sim 0.4$. Meanwhile in the B-ring the troughs yield a photometric optical depth of $\sim 0.8$. It is true that self-gravity wakes complicate the relationship between dynamical and photometric optical depth, yet the discrepancy is concerning. Second, is the observed mean optical depth differs in the wave and in the flat zones: in the former it is approximately 0.8; in the latter is is closer to 1.3. This further complicates the mapping of our results to the observations, and indicates our theory requires additional refinement. Comparison of our solutions to C-ring observations must first resolve one key question: does the free evolution of the BTI generate the low-amplitude 1000-km undulations, found between 77,000 km and 86,000 km, or the larger-amplitude 100-km plateaus, between 84,000 and 91,000 km (Fig.~13.17 in Colwell et al.~2009)? On account of the small amplitudes and morphology of our wavetrain solutions, we conclude that the 1000-km undulations are the result of the BTI working alone. The plateaus are probably caused by something else, though the ballistic transport process may influence their general shape (Estrada \& Durisen 2010). If both the 1000-km undulations in the C-ring and the 100-km waves in the B-ring are BTI wavetrains then it follows that $l_\text{th}$ could vary significantly between the two radial locations. This variation may arise from differences in the sizes, composition, or regolith properties of the ring particles, on the one hand, or the trajectories and speeds of the incoming meteoroids, on the other. For example, recent spectroscopic studies indicate that the sizes of regolith grains vary with radius (Morishima et al.~2012, Hedman et al.~2013). But it is unclear whether this means particles are more or less `fluffy' (and hence $l_\text{th}$ smaller or greater) in different ring regions. We view this as a key question in the study of ballistic transport, deserving of further study\footnote{Note that the recent impacts observed by Tiscareno et al.~(2013) involved cm to m sized meteoroids and, being in a different collisional regime, cannot help constrain $l_\text{th}$.}. In our following paper, the role of these invariant solutions is made clear through full time-dependent simulations. There we also run a suite of simulations of the inner B-ring edge, which itself could be unstable to the BTI. Further work will improve our basic model, through the addition of more physical processes. For instance, the ring's viscosity should be an increasing function of $\tau$, not a constant as assumed here. Preliminary results, however, show no qualitative changes arises from this effect. Of greater importance may be the form of the absorption probability, $P$. Throughout this paper, we assume it only depends on the absorbing radius. But in lower optical depth regions it will also depend on the ejecta emitting radius. This effect will influence both the B and C-rings, the former on account of the low $\tau$ achieved in wavetrain troughs. | 14 | 4 | 1404.4245 |
1404 | 1404.5305_arXiv.txt | We present the first quantitative UV spectroscopic analysis of resolved OB stars in IC~1613. Because of its alleged very low metallicity ($\lesssim$1/10 \Zsun, from \ion{H}{2} regions), studies in this Local Group dwarf galaxy could become a significant step forward from the SMC towards the extremely metal-poor massive stars of the early Universe. We present HST-COS data covering the $\sim$1150-1800\AA~ wavelength range with resolution R$\sim$2500. We find that the targets do exhibit wind features, and these are similar in strength to SMC stars. Wind terminal velocities were derived from the observed P~Cygni profiles with the SEI method. The \vinf-Z relationship has been revisited. The terminal velocity of IC~1613 O-stars is clearly lower than Milky Way counterparts, but there is no clear difference between IC~1613 and SMC or LMC analogue stars. We find no clear segregation with host galaxy in the terminal velocities of B-supergiants, nor in the \vinf/\vesc~ ratio of the whole OB star sample in any of the studied galaxies. Finally, we present first evidence that the Fe-abundance of IC~1613 OB stars is similar to the SMC, in agreement with previous results on red supergiants. With the confirmed $\sim$1/10 solar oxygen abundances of B-supergiants, our results indicate that IC~1613's $\alpha$/Fe ratio is sub-solar. | \label{s:intro} The great observatories of this and the coming decade (ALMA, JWST, and E-ELT) will bring direct observations of the epoch of re-ionization. In the infant Cosmos the first, very massive stars played a crucial role as ionizing sources \citep{Ral10} and likely progenitors to long-GRBs \citep{Gal09,WH06}. An accurate theoretical framework for the evolution of metal-free massive stars, and subsequent calculations of stellar yields and ionizing power, are key to interpret the coming observations. Radiation-driven winds are one of the main pillars of said theory, as agents of mass and momentum removal during the evolution of massive stars. So far the models for the high-redshift, metal-poor Universe rely on theoretical predictions, or on stellar libraries of the Small Magellanic Cloud (SMC). However, recent results % indicate that the winds of sub-SMC metallicity massive stars may differ significantly from their higher metallicity counterparts. If this is confirmed, massive star evolution and feedback would need deep revision, which would then need to be propagated to early Universe models. In the classical radiation-driven wind framework the wind mass loss rate (\Mdot) decreases with decreasing metallicity \citep{VKL00,VKL01,PSL00}, a prediction confirmed by spectroscopic studies of OB stars down to SMC metallicities \citep[e.g.][]{Mal07b}. The scaling relation is $\rm \dot M \propto Z^{\alpha}$, with $ \rm \alpha \sim 0.7-0.8$ \citep{VKL01,Mal07b}. The dependence of the wind terminal velocity on metallicity has only been established empirically and is allegedly much weaker, $\rm v_{\infty} \propto Z^{0.13}$~ \citep{LRD92}. This relation has not been reviewed recently, yet it was used by \citet{VKL01} and \citet{Mal07b} to derive the \Mdot-Z dependence (see below). At extremely low metallicities ($\rm Z \leq 0.01 Z_{\odot}$), \Mdot~ and \vinf~ are expected to decrease with decreasing metallicity more steeply \citep{K02}. By contrast, very metal poor Luminous Blue Variable stars with strong optical P~Cygni profiles have been found in the Local Group \citep{Hal10} and in farther galaxies \citep{DCS01,pustilnik08,izotov11}. \citet{Tal11} report six O-type stars with stronger wind momentum than expected in Local Group galaxies with metallicity $\rm \sim 1/7-1/10 \, Z_{\odot}$. % We studied an Of star in IC\,1613 possibly exhibiting a strong wind or, alternatively, anomalous wind acceleration \citep{Hal12}. These examples challenge the currently accepted metallicity scaling relations for \Mdot, used by stellar evolution models \citep{MM00,HL00} and ultimately galactic chemical evolution models (see \citealt{HR10}). While the analysis with more complete models may explain these findings (see discussion by \citealt{L12}), they may constitute the first hints for a more complex wind-driving mechanism. Better observations on extended samples are necessary to establish or refute this apparent contradiction to the theory and, in particular, ultraviolet (UV) spectroscopy covering at least the $\sim$1150-1800\AA~ range is key. The installation of the Cosmic Origins Spectrograph on-board the Hubble Space Telescope (HST-COS) has enabled UV spectroscopy of sources a factor of $\sim$ 10 fainter than before. Access has been enabled to metal-poorer galaxies than the SMC, which are inconveniently located farther in the Local Group, but whose stellar population can still be resolved with the HST \citep{NUVA}. Prior to COS, their winds could only be studied from optical spectral lines (H$_{\alpha}$~ and \ion{He}{2}$\lambda$4686), invariant to the wind velocity law (parameterized by its exponent $\beta$) and the terminal velocity unless the wind is very strong. In fact, the synthetic spectra for these transitions exhibit a degeneracy to a combination of parameters involving \vinf, \Mdot, and the stellar radius \Rstar~ ($ Q = \dot M / (v_{\infty} \cdot R_{\ast})^{1.5}$, \citet[][eq. 2]{KP00}). In addition, the uncertainties of $\beta$~ translate into significant uncertainties in $Q$ (hence \Mdot). Both \Mdot~ and \vinf~ are required to calculate the modified wind-momentum ($D_{mom} = \dot M \cdot v_{\infty} \cdot R_{\ast}^{0.5}$, or equivalently $D_{mom} = Q \cdot v_{\infty}^{2.5} \cdot R_{\ast}^{2}$). The correlation between the stellar wind-momentum and luminosity \citep[the WLR,][]{Kal95} is our most powerful tool to evaluate radiation-driven wind strength. Lacking diagnostics for the terminal velocity, optical studies set \vinf~ from empirical relations to the escape velocity which exhibit large scatter \citep[e.g. \vinf/\vesc=2.65 for \Teff$\geq$21000~K,][]{KP00} and then scale it with metallicity using Leitherer's relation. This translates into large uncertainties in the actual mass loss rate of the stars, calculated from $Q$, and the stellar wind momentum. In particular, the errors of \vinf~ propagate to $D_{mom}$~ so that $\Delta (\log D_{mom}) = 2.5 \Delta (\log v_{\infty})$, for $Q$ and \Rstar~ constant. We present first results of our program to study the winds of OB-stars in the metal-poor galaxy IC~1613 with HST-COS UV spectroscopy. Our team is thoroughly characterizing the population of blue massive stars in this Local Group galaxy with very low metal content \citep[0.13 \Zsun~ from \ion{H}{2} regions,][]{Fal07}. IC~1613 is one of the closest dIrr to the Milky Way \citep[DM=24.27,][]{Dal01}, with low foreground extinction \citep[E(B-V)=0.02,][]{LFM93}. Thanks to COS enhanced sensitivity, we can perform ultraviolet spectroscopy of the targets, and obtain unique insight into their winds. The wind stratification will be studied from the P~Cygni profiles of metallic lines, providing information on \Mdot, $\beta$~ and \vinf. The degeneracy of optical synthetic spectra to the $Q$ parameter will be thus broken, and more accurate values will be input into the WLR to assess the winds. The paper is organized as follows. The HST-COS observations are described in Section~\ref{s:obs}, and concerns on data reduction in Section~\ref{s:red}. The morphology of the observed spectra is discussed in Section~\ref{s:mor} and compared to LMC and SMC counterparts. Section~\ref{s:sei} details how the terminal velocities were derived, with results presented in Section~\ref{s:r}. They allow us to evaluate how \vinf~ changes in Local Group galaxies in Section~\ref{s:vinfZ}. We reflect upon IC~1613 metal content in Section~\ref{s:metal}, and provide our final conclusions in Section~\ref{s:fin}. | 14 | 4 | 1404.5305 |
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1404 | 1404.5133_arXiv.txt | We estimate the amount of angular momentum transferred by the low-frequency oscillations detected in the rapidly rotating {hot} Be star HD\,51452. Here, we assume that the oscillations detected are stochastically excited by convective motions in the convective core of the star, that is, we treat the oscillations as forced oscillations excited by the periodic convective motions of the core fluids having the frequencies observationally determined. With the observational amplitudes of the photometric variations, we determine the oscillation amplitudes, which makes it possible to estimate the net amount of angular momentum transferred by the oscillations using the wave-meanflow interaction theory. Since we do not have any information concerning the azimuthal wavenumber $m$ and spherical harmonic degree $l$ for each of the oscillations, % {we assume} that all the frequencies detected are prograde or retrograde in the observer's frame and they are all associated with a single value of $m$ both for even modes ($l=|m|$) and for odd modes ($l=|m|+1$). We estimate the amount of angular momentum transferred by the oscillations for $|m|=1$ and 2{, which are typical $|m|$ values for Be stars,} and find that the amount is large enough for a decretion disc to form around the star. % {Therefore, transport of angular momentum by waves stochastically excited in the core of Be stars might be responsible for the Be phenomenon.} | Be stars are rapidly rotating {active} late O, B, or early A stars hosting a circumstellar decretion disc {fed} by discrete mass loss events. {The} mass ejections and {disc} produce emission lines in the optical spectrum of Be stars (see, e.g., Porter \& Rivinius 2003; Rivinius et al. 2013 for recent reviews on Be stars), and {it is believed} that the circumstellar discs around Be stars are viscous Keplerian discs (Lee, Saio {\&} Osaki 1991). Be stars of late O and early B type are also known as $p$-mode pulsators, while those of late B and early A type are $g$-mode pulsators, where both the $p$-mode and $g$-mode pulsations are excited by the opacity mechanism associated with the opacity bump produced by iron-peak elements at the temperature regions of $T\sim 2\times10^5$K in their envelope (e.g., Pamyatnykh 1999). { In addition, the convective core of intermediate-mass and massive stars are able to {stochastically excite} oscillation modes (e.g., Belkacem, Dupret, Noels 2010; Samadi et al.. 2010, Shiode et al. 2013). {Propagative gravity (gravito-inertial in the case of rapid rotation) waves and modes are excited (e.g., Browning, Brun, Toomre 2004; Rogers et al. 2013; Mathis, Neiner \& Tran Minh 2014), because of the convective motions and of their penetration into the surrounding radiative envelope.} These waves are able to transport angular momentum because of their dissipation along their propagation and at corotation layers ({see} e.g., Zahn, Talon, Matias 1997 and Alvan, Mathis, Decressin 2013 for gravity {waves,} and Pantillon et al. 2007 and Mathis et al. 2008 for gravito-inertial waves).} Mechanisms for the mass ejections and disc formation, however, have not yet been identified for Be stars. For disc formation mechanisms, {several models have been proposed} such as stellar wind models (e.g., Bjorkman \& Cassinelli 1993; Owocki, Cranmer, \& Blondin 1994; Cranmer \& Owocki 1995), a model making use of magnetic fields to support a disc (Cassinelli et al. 2002; Owocki \& ud Doula 2002), an evolutionary model that assumes angular momentum redistribution in the interior of rotating stars (e.g., Ekstr\"om et al. 2008; Granada et al. 2013), and a model using {$\kappa$-driven} pulsations as a carrier of angular momentum to the surface region of rapidly rotating stars (e.g., Ando 1983, 1986; Lee \& Saio 1993; Cranmer 2005, 2009; Ishimatsu \& Shibahashi 2013). The stellar wind models are not necessarily successful in producing Keplerian discs around the stars. Magnetic fields strong enough to support circumstellar discs are not common among Be stars { (e.g., Neiner et al. 2012b)}. On the other hand, { as discussed above}, stellar pulsations in Be stars are excited by the opacity bump mechanism or{/and} probably by stochastic excitation in the convective core. In this sense using the pulsations as the angular momentum carrier can be a promising mechanism for disc formation around Be stars, particularly when both stellar evolution and pulsation play their own parts. In fact, as suggested by Lee (2013), steady viscous decretion disc solutions are possible if a sufficient amount of angular momentum is deposited {in} the surface layers. Recently, Neiner et al. ({2012a}) observed {the} hot Be star HD\,51452 {with} the CoRoT {satellite} and identified numerous low-frequency oscillations. Since the star is rapidly rotating at the rate 1.22 cycle per day and the oscillation frequencies in the observer's frame are comparable to or {lower} than the rotation frequency, the oscillations are low-frequency ones in the corotating frame of the star, suggesting that the effects of rapid rotation of the star can be significant to determine the wave properties propagating in the star. It is also important to note that low-frequency oscillations in early B type main sequence stars such as HD\,51452 cannot be excited by the opacity bump mechanism. Neiner et al. ({2012a}) therefore suggested that the low-frequency oscillations detected are {gravito-inertial waves} excited stochastically by the convective motions in the core of the star. In this paper, we estimate a possible amount of angular momentum transferred by the low-frequency oscillations identified by Neiner et al. ({2012a}) for the Be star HD\,51452. We calculate non-adiabatic low-frequency oscillations by taking the effects of rotation of the star {into account} assuming that the rotation is uniform (Lee \& Baraffe 1996), and we employ a wave-meanflow interaction theory (e.g., Andrews \& McIntyre 1978ab; Grimshaw 1984; Lee 2013) to estimate the amount of angular momentum transfer. The oscillations are treated as forced oscillations whose frequencies are equal to those identified {with CoRoT} for the star. Since the CoRoT {observations do} not provide any information concerning the azimuthal wavenumber $m$ and spherical harmonic degree $l$ for the oscillations, we employ in this paper a working hypothesis that all the oscillations identified have the same $m$ and $l$ when estimating the possible amount of angular momentum. We describe the {analytical} method of calculation in \S2 and numerical results in \S 3. Conclusions are given in \S 4. { | } Using a theory of wave-meanflow interaction, we have estimated the possible amount of angular momentum transferred by {gravito-inertial waves} having the set of frequencies observationally detected for {the} Be star {HD\,51452}. Since the Be star is rapidly rotating and the detected frequencies are low in the observer's frame in the sense that $\sigma\la\Omega_{\rm crit}$, the frequencies $\omega$ in the corotating frame of the star {are} low, that is, the detected frequencies are in the frequency domain of $g$-modes and $r$-modes. Since the opacity {$\kappa$} mechanism does not work for excitation of low frequency modes in massive stars of $M_*\ga 10M_\odot$ {for a solar metallicity (Pamyatnykh 1999, Miglio et al. 2007)}, we need to look for an alternative mechanism for the excitation of the observationally detected low frequency oscillations. In this paper, we assume that the oscillations are excited by periodic convective motions of the core fluid in the massive star {as suggested by Neiner et al. (2012a)}. We treat therefore the detected oscillation frequencies as forcing frequencies for waves propagating in the radiative envelope. We calculate the forced oscillations having the observed frequencies by imposing pressure perturbations produced by the convective fluid motions {at the core-envelope interface as the boundary condition}. This procedure makes inhomogeneous the system of linear differential equations for oscillations, which can be integrated for arbitrary frequencies. In the theory of wave-meanflow interaction, {the high-$\kappa$} regions works for accelerating (decelerating) the rotation (mean)flows for retrograde (prograde) waves observed in the corotating frame of the star. Since we have no information {on} the azimuthal wave number $m$ for the detected oscillations, we calculated the oscillations of the frequencies just assuming that they are all retrograde modes or prograde modes in the observer's frame. {In} the former case, all the modes are retrograde in the corotating frame and hence the angular momentum deposition can take place efficiently in the surface layers of the star. {In} the latter case, however, the modes can be separated into prograde and retrograde modes, which means that even in {the high-$\kappa$} regions the net amount of angular momentum deposition depends on the net sum of accelerating and decelerating contributions of retrograde and prograde waves. {Imposing the observed} frequencies, we have shown that the amount of angular momentum transferred to the surface regions is large enough for decretion disc formation with the mass loss rate $\sim10^{-10}M_\odot/\rm yr$, although a definite amount of angular momentum transferred can be estimated only after the details {of mode identification are determined. This could possibly be done with a ground-based high-resolution spectroscopic campaign.} Since we assume forced oscillations propagating in the radiative envelope of the star, there exist the possibility that some of the forcing frequencies, { that are determined by properties of motions in the convective core}, are in resonance with {eigen}frequencies of the {free} oscillation modes such as $g$-modes or $r$-modes in the envelope. Both the forcing frequencies and the {eigen}frequencies can change with {the} {star's} evolution. If {such} a frequency resonance {happens} between an {envelope} {free} mode and a forcing frequency, {its} {amplitude} can be increased so that the effects of angular momentum deposition or extraction in the surface regions is enhanced. {In this paper, we have demonstrated, using the case of HD\,51452, the ability of gravito-inertial waves to transport angular momentum efficiently {in} rapidly rotating massive stars} {and to deposit this momentum just below the surface. This mechanism increases the velocity of the surface layers. These layers can then reach the escape velocity at which material gets ejected from the star. Therefore, this mechanism could be at the origin of matter ejections in Be stars and the reason of the presence of a decretion disc around Be stars.} {In a near future, we will improve the physical modelling of this wave-meanflow. First, in this work, we have assumed a uniform rotation to compute the oscillations. However, because angular momentum deposition occurs in a rather short time-scale in the surface layers, it may be important to take {differential rotation into account} when computing modes and the transport of angular momentum they {induce} (e.g., Lee \& Saio 1993; Mathis 2009). In this way, we will get a complete picture of wave-meanflow interactions. Next, the modelling of the amplitude of waves, which are stochastically excited by the convective core in massive stars, must be improved using theoretical models (e.g., Belkacem et al. 2009; Lecoanet \& Quataert 2013; Mathis, Neiner \& Tran Minh 2014) and {more and more realistic} direct numerical simulations (e.g. Browning et al. 2004, Rogers et al. 2013). Besides, differential rotation and related meridional flows and turbulence play an important role in the evolution of massive stars (e.g., Meynet \& Maeder 2000). In this context, it would be important to take these mechanisms {into account as they} also transport angular momentum simultaneously with gravito-inertial waves. This has {already} been done for the case of solar-type stars (Talon \& Charbonnel 2005, Charbonnel et al. 2013, Mathis et al. 2013), in which waves strongly modify the star's rotational evolution, {but} the case of massive stars {still has} to be studied. Such study {should} be undertaken first in spherical models {in which} the centrifugal acceleration is treated as a perturbation to unravel the role of each physical processes. Then, complete 2D models must be developped for rapidly rotating stars (e.g., Ballot et al. 2010; Espinosa Lara \& Rieutord 2013).} \bigskip | 14 | 4 | 1404.5133 |
1404 | 1404.7136_arXiv.txt | Fomalhaut is a triple system, with all components widely separated ($\sim 10^5$~au). Such widely separated binaries are thought to form during cluster dissolution, but that process is unlikely to form such a triple system. We explore an alternative scenario, where A and C form as a tighter binary from a single molecular cloud core (with semimajor axis $\sim 10^4$~au), and B is captured during cluster dispersal. We use N-body simulations augmented with the Galactic tidal forces to show that such a system naturally evolves into a Fomalhaut-like system in about half of cases, on a timescale compatible with the age of Fomalhaut. From initial non-interacting orbits, Galactic tides drive cycles in B's eccentricity that lead to a close encounter with C. After several close encounters, typically lasting tens of millions of years, one of the stars is ejected. The Fomalhaut-like case with both components at large separations is almost invariably a precursor to the ejection of one component, most commonly Fomalhaut C. By including circumstellar debris in a subset of the simulations, we also show that such an evolution usually does not disrupt the coherently eccentric debris disk around Fomalhaut A, and in some cases can even produce such a disk. We also find that the final eccentricity of the disk around A and the disk around C are correlated, which may indicate that the dynamics of the three stars stirred C's disk, explaining its unusual brightness. | \label{section:intro} Fomalhaut has been suspected to be part of a widely separated binary star system for some time now \citep{1938AJ.....47..115L}. Recent analysis has confirmed that the K4V star TW PsA has both a similar proper motion and radial velocity to Fomalhaut, such that it is not an interloping field star, but forms a binary star system with Fomalhaut. The pair have a three-dimensional separation of $5.74^{+0.04}_{-0.03}\times 10^4 \rm{au}$. Combining isochronal, rotational, X-ray, and lithium ages for the pair, the system is constrained to have an age of $440 \pm 40~\rm{Myrs}$, the mass of Fomalhaut A to be $1.92 \pm 0.02 M_{\oplus}$, and the mass of Fomalhaut B to be $0.73^{+0.02}_{-0.01}M_{\oplus}$ \citep{2012ApJ...754L..20M}. More recently, a third member of the Fomalhaut system has been recognised. The M4V star LP 876-10, or Fomalhaut C, is a $0.18 \pm 0.02 M_{\odot}$~star with a three dimensional separation from Fomalhaut A of $1.58^{+0.02}_{-0.01}\times 10^5 \rm{au}$~\citep{2013AJ....146..154M}, which also has a common proper motion with Fomalhaut A and B. The uncertainty in the measured velocities of the three stars is roughly half a kilometre per second, much smaller than the relative velocities expected between field stars, allowing the inference that the system is bound. However, the escape speed for B is about a quarter kilometre per second, and C's smaller still, so little can be said about the orbital configuration. The dynamics of the Fomalhaut system are particularly interesting as Fomalhaut A is known to harbour a debris disk \citep{1986ASSL..124...61G}, which has a coherent eccentricity of $0.11 \pm 0.01$ \citep{2005Natur.435.1067K}. The origin of the debris disk's eccentricity is not known. It has been suggested it may result from the action of one or more shepherding planet(s) \citep{2006MNRAS.372L..14Q,2009ApJ...693..734C,2012ApJ...750L..21B}, although other possible origins have been advanced \citep[e.g.,][]{2013Natur.499..184L}. Fomalhaut A is also accompanied by a point-like object dubbed 'Fomalhaut b' \citep{2008Sci...322.1345K}. Non-detection of Fomalhaut b at thermal wavelengths led \citet{2008Sci...322.1345K} to the suggestion that it may be the dust cloud ejected in a collision between planetesimals, or a circumplanetary ring, and \citet{2011MNRAS.412.2137K} modelled it as a circumplanetary dust cloud created by the collisions of irregular satellites. Continuing observation has allowed for improved fits to the orbit of Fomalhaut b, which have revealed that it is not a shepherding planet, but has an eccentricity of $\sim 0.8$, and a semimajor axis similar to the dust belt bodies \citep{2013ApJ...775...56K,2014A&A...561A..43B}. This eccentric orbit precludes Fomalhaut b having a mass significantly above $10 M_{\oplus}$, unless it was recently scattered to its highly eccentric orbit, as a more massive planet would perturb the disk to disruption \citep{2014A&A...561A..43B,2014MNRAS.tmp..149T}. The origin of the eccentricities of the dust belt's eccentricity, Fomalhaut b's eccentricity, and the nature of Fomalhaut b all remain open questions. Moreover, Fomalhaut C has also been discovered to have a dust disk \citep{2014MNRAS.438L..96K}. Debris disks are rare around M stars \citep{2007ApJ...667..527G}, and thus in the Fomalhaut system, which offers the possibility of additional dynamical constraints, the existence of a debris disk about an M star is of particular interest. The formation of such a wide triple system presents some problems. The separations of both components are larger than the typical sizes of star forming cores and thus, given that such cores rotate at far less than break-up velocity \citep{1993ApJ...406..528G}, the system contains far too much angular momentum to have been created by conventional core fragmentation. The unfolding of triple stellar systems suggested by \citep{2012Natur.492..221R} appears to be inapplicable to Fomalhaut, which lacks an inner binary. Capture during cluster dispersal has also been suggested as a mechanism for wide binary formation \citep{2010MNRAS.404.1835K,2011MNRAS.415.1179M}, but the low probability of capture makes the formation of a triple system by this method highly unlikely. The fraction of field stars with wide binaries is $\sim 10^{-2}$, \citep{2010AJ....139.2566D}, suggesting that the chance of forming a cluster capture binary $\sim 10^{-2}$, and thus forming a triple this way should have odds $\sim 10^{-4}$. In comparison, the chance of forming a $10^4$~au binary is $\sim 10^{-1}$, \citep{2014MNRAS.437.1216D}. Thus we are motivated to consider an alternative scenario, where AC form as a tighter binary from a single core, and C is moved outwards by interactions with B, which is captured as a wide binary. We describe our numerical method in \textsection \ref{section:method}, and present the results of our simulations in \textsection \ref{section:results}. We perform a subset of simulations with debris disks encircling Fomalhaut A and C in \textsection \ref{section:disk} to assess the compatibility of this scenario with the observed disks. | Formation of a weakly bound triple system like Fomalhaut is unlikely by capture of two stars during cluster dispersal. We consider an alternate scenario, where Fomalhaut A and C formed as a tighter binary (with semimajor axis $5000~\rm{au} \lesssim a \lesssim 13600~\rm{au}$)~and Fomalhaut B was captured into a weakly bound orbit during cluster disperal. We simulate the evolution of such a system, and show that such a system commonly ($55\% - 60\%$) evolves into a Fomalhaut-like system, with both components at large separations from Fomalhaut A, on timescales compatible with the current age of the system. In this evolution, the present day state where both components are on wide orbits is a temporary one, which typically dissociates in $\mathcal{O}\left(10\right)$~Myrs. In our simulations, systems that match Fomalhaut may be bound, or else be in the process of ejecting C from the system. % We perform an additional 50 simulations with disks of test particles around Fomalhaut A and Fomalhaut C, to ascertain whether this formation scenario is compatible with the coherent eccentric disk around Fomalhaut A, and the existence of a detectable disk around Fomalhaut C. Nineteen of those simulations became matches for the Fomalhaut system. In seven cases, the disk around Fomalhaut A is not significantly perturbed, and such cases would be compatible with the dynamics of A's disk being set by an internal source \citep[see e.g.,][]{2006MNRAS.372L..14Q}. Intriguingly, five systems develop coherent eccentricities in A's disk, owing to secular interactions or close encounters with Fomalhaut C in its much tighter primordial orbit about A, with eccentricities with a factor of 5 of the present day value. This suggests a possible origin of A's coherently eccentric disk. We also find that the final excitation of disk eccentricity around A and around C are correlated, suggesting that if A's debris disk was driven to high eccentricity by interactions with C, C's disk may be expected to have been driven to high eccentricity by A. This may explain why Fomalhaut C has a detectable debris disk, which is rarely found around an M star. The total likelihood of the scenario presented may seem quite small, when one considers that Fomalhaut is the fourth nearest A star. However, in matching the details of any system, and focussing on the most interesting aspects, the intrinsic likelihood of that outcome will be small. The probability that the system we conjectured to be the primordial state of Fomalhaut undergoes an instability to produce a Fomalhaut-like system is $\sim 50\%$. Although the probability of forming our conjectured initial state is small, $10^5 \rm{au}$~binaries are an order of magnitude less common than $10^4 \rm{au}$ binaries \citep{2010AJ....139.2566D,2014MNRAS.437.1216D} , and thus given the high chance of creating a Fomalhaut like system from an initially hierarchical system, this should be the preferred model. There are two observational tests of this scenario. Firstly (as noted above) we find a high likelihood ($> 80\%$) that detailed orbital characterisation of the Fomalhaut system will reveal that C's orbit is such that it interacts strongly with B during pericentre passage. Secondly, further circumstantial evidence in favour of this scenario would be provided if C were found to possess an eccentric disc. However, this too is not an inevitable consequence of this scenario: Figure \ref{fig:diskcomp} demonstrates that there are situations where A's disc acquires an eccentricity of $\gtrsim 0.1$~while the eccentricity of C's disk remains substantially smaller. | 14 | 4 | 1404.7136 |
1404 | 1404.5902_arXiv.txt | {Results from $UBVRI$ optical photometric observations of the pre-main sequence star V350 Cep during the period 2004$-$2014 are presented. The star was discovered in 1977 due to its remarkable increase in brightness by more than 5 mag ($R$). In previous studies, V350 Cep was considered a to be a potential FUor or EXor eruptive variable. Our data suggest that during the period of observations the star maintains its maximum brightness with low amplitude photometric variations. Our conclusion is that V350 Cep was probably an intermediate object between FUors and EXors, similar to V1647 Ori. | % \label{sect:intro} Studies of pre-main sequence (PMS) stars are very important for modern astronomy because they give an opportunity to understand the early stages of stellar evolution, as well as to test stellar evolution scenarios. Depending on their initial mass, young stars pass through different periods of stellar activity. The most prominent manifestations of this activity are changes in the star's brightness with various periods and amplitudes. Photometric and spectroscopic variabilities are the most common characteristics of PMS stars. The most widely-spread type of PMS objects $-$ T Tauri stars are young, low mass stars ($M \leq 2M_\odot$). Their study began after the pioneering work of \cite{Joy+1945}. The main characteristics of T Tauri stars are their emission spectra and their irregular photometric variability. Some T Tauri stars exhibit strong brightness variations over comparatively short time intervals (days, months) with amplitudes of up to several magnitudes. T Tauri stars are separated in two subclasses: classical T Tauri stars (CTTS) surrounded by massive accreting circumstellar disks and weak-line T Tauri stars (WTTS) without evidence for disk accretion (\citealt{Bertout+1989}). \newpage According to \cite{Herbst+etal+1994} photometric variability of WTTS is due to the rotation of the stellar surface covered with large cool spots. The periods of variability in WTTS are observed on time scales of days and with amplitudes up to 0.8 mag in the $V$-band. Variability of CTTS is more complicated: the variability is caused by a superposition of cool and hot surface spots producing non-periodic variations with amplitudes up to 2-3 mag. in the $V$-band. The large amplitude outbursts of PMS stars can be grouped into two main types, named after their respective prototypes: FU Orionis (FUor; \citealt{Ambartsumian+1971}) and EX Lupi (EXor; \citealt{Herbig+1989}). Both types of stars are probably related to low-mass T Tauri stars with massive circumstellar disks, and their outbursts are generally attributed to a sizable increase in accretion rate from the circumstellar disk onto the stellar surface. The outburst of FUor objects last for several decades, and the rise time is shorter than the decline. EXor objects show frequent (every few years or a decade), irregular or relatively brief (a few months to one year) outbursts with an amplitude of several magnitudes ($\Delta$$V$$\approx$3-5). The PMS star V350 Cep is located in the field of the reflection nebula NGC 7129, a region with active star formation. The region is immersed in a very active and complex molecular cloud (\citealt{Hartigan+Lada+1985}; \citealt{Miranda+etal+1993}). The distance to NGC 7129 as determined by \cite{Straizys+etal+2014} is 1.15 kpc. Variability in V350 Cep was discovered by \cite{Gyulbudaghian+Sarkissian+1977} who compared their photographic observations of NGC 7129 with the Palomar Observatory Sky Survey (POSS) plates. V350 Cep was not seen on the POSS O-plate obtained in 1954 (limit $\sim$ 21 mag) and is slightly above the limit of the E-plate. The measured brightness of the star in 1977 was approximately 17.5 mag in $B$-band and 16.5 mag in $V$-band. Follow-up observations by the Sternberg Astronomical Institute (\citealt{Pogosyants+1991}), Sonneberg and Alma-Ata Observatories, which are in plate archives (\citealt{Gyulbudaghian+1980}), suggest that V350 Cep was below the plate limits before 1970, i.e. it was fainter than 17.5 mag in $B$-band. The spectral class of V350 Cep is defined as M2 by \cite{Cohen+Fuller+1985} and as M0 by \citealt{Kun+etal+2009}. Photometric observations of V350 Cep (\citealt{Gyulbudaghian+Sarkissian+1978}; \citealt{Hakverdian+Gyulbudaghian+1978}; \citealt{Shevchenko+Yakubov+1989}; \citealt{Pogosyants+1991}; \citealt{Semkov+1993}, \citealt{Semkov+1996}, \citealt{Semkov+1997}, \citealt{Semkov+2002}, \citealt{Semkov+2004a}; \citealt{Semkov+etal+1999}) demonstrated changes of brightness, which are typical for CTTS with an amplitude of about 1.5 mag in the $B$-band. All spectral observations of V350 Cep (\citealt{Gyulbudaghian+etal+1978}; \citealt{Magakian+Amirkhanian+1979}; \citealt{Cohen+Fuller+1985}; \citealt{Goodrich+1986}; \citealt{Miranda+etal+1994}; \citealt{Magakian+etal+1999}; \citealt{Semkov+2004b}; \citealt{Kun+etal+2009}) suggest that its spectrum is similar to the CTTS spectra, including being quite variable, having an emission spectrum and a variable P Cygni profile fot the H$\alpha$ line. Collected photometric data indicate that the rise in brightness began some time before 1970, and the light curve of the star resembles that of a classic FUor star V1515 Cyg (see \citealt{Clarke+etal+2005}). Section 2 gives information about telescopes and cameras used and data reduction. Section 3 describes the derived results and their interpretation. \newpage | The presented photometric data are a continuation of our long-term photometric study of V350 Cep. The $UBVRI$ lights curves of V350 Cep from all our CCD observations (\citealt{Semkov+1996}, \citealt{Semkov+1997}, \citealt{Semkov+2002}, \citealt{Semkov+2004a}; \citealt{Semkov+etal+1999} and the present paper) are shown in Figure~\ref{Fig1}. In the figure, circles denote CCD photometric data acquired with the 2-m RCC telescope; triangles $-$ the photometric data taken with the 1.3-m RC telescope; diamonds $-$ the photometric data collected with the 50/70-cm Schmidt telescope, and squares $-$ the photometric data obtained with the 60-cm Cassegrain telescope. \begin{figure} \centering \includegraphics[width=12.0cm, angle=0]{ms1799fig1.eps} \caption{CCD $UBVRI$ light curves of V350 Cep for the period August 1993 $-$ February 2014} \label{Fig1} \end{figure} The data reported in the present paper indicate that the brightness of V350 Cep remained close to the maximum value during the period 2004$-$2014 (Table~\ref{Tab1}). Thus, the star has been keeping its maximum brightness during the past 35 yr and for the same period it showed photometric variability with a low amplitude. The observed amplitudes in the period 1993$-$2014 are 0.47 mag for the $I$-band, 0.87 mag for the $R$-band, 1.19 mag for the $V$-band, 1.30 mag for the $B$-band and 0.84 mag for the $U$-band. These values are typical of T Tauri stars surrounded with an accreting circumstellar disk. Figure~\ref{Fig2} shows the long-term $B/pg$-light curve of V350 Cep from all available observations. The circles denote our CCD photometric data (\citealt{Semkov+1996}, \citealt{Semkov+1997}, \citealt{Semkov+2002}, \citealt{Semkov+2004a}; \citealt{Semkov+etal+1999} and the present paper); triangles $-$ the photographic data from the Rozhen Schmidt telescope (\citealt{Semkov+1993}; \citealt{Semkov+1996}); diamonds $-$ the photographic data from \cite{Pogosyants+1991}; squares $-$ the photographic data from the Asiago Schmidt telescope (\citealt{Semkov+etal+1999}); empty diamonds symbols $-$ the photographic data from \cite{Shevchenko+Yakubov+1989}; pluses $-$ the photographic data from Byurakan Schmidt telescope (\citealt{Gyulbudaghian+Sarkissian+1977}; \citealt{Semkov+1993}); the empty triangles $-$ the limit of the photographic data from the POSS plates, the Sternberg Astronomical Institute plate archive \cite{Pogosyants+1991} and the Asiago Schmidt telescope plate archive (\citealt{Semkov+etal+1999}). The available photometric data suggest that the period of strong increase in brightness continued to about 1978 was followed by a period of irregular variability around the level of maximum brightness lasting up to now. \begin{figure} \centering \includegraphics[width=12.0cm, angle=0]{ms1799fig2.eps} \caption{$B/pg$ light curve of V350 Cep from all available observations} \label{Fig2} \end{figure} \newpage Another important result from our photometric study is the variation of color indices with stellar brightness. The measured color index $V-I$ versus the stellar magnitude $V$ during the period of our CCD observations is plotted in Figure~\ref{Fig3}. A clear dependence can be seen from the figures: the star becomes redder as it fades. The other indices $V-R$ and $B-V$ show a similar trend on the color-magnitude diagram. Such color variations are typical of T Tauri stars with large cool spots, whose variability is produced by rotation of the spotted surface. Consequently, V350 Cep shows photometric characteristics of WTTS (variability with small amplitude in a time scale of days). On the other hand, the observed spectra of V350 Cep can be classified as a CTTS spectrum (\citealt{Magakian+etal+1999}; \citealt{Semkov+2004b}). As can be seen from the Table~\ref{Tab1}, V350 Cep shows a very strong ultraviolet excess $-$ a characteristic also typical of CTTS. Moreover, the long-term light curve of V350 Cep is similar to FUor type objects such as V1515 Cyg. These discrepancies make V350 Cep a unique object, which is very difficult to classify. \begin{figure} \centering \includegraphics[width=10.0cm, angle=0]{ms1799fig3.eps} \caption{Relationship between $V$ magnitude and $V-I$ color index in the period of all our CCD observations} \label{Fig3} \end{figure} Regardless of its similarity to T Tauri stars, the large amplitude outburst of V350 Cep can only be explained only as an episode of enhanced accretion. We suggest that V350 Cep is an object similar to V1647 Ori (see \cite{Aspin+Reipurth+2009} and references therein). The eruptive PMS star V1647 Ori was discovered in 2004 during its large amplitude outburst and is considered to be a unique object which shows the photometric characteristics of FUors and spectral characteristics of EXors. Both stars V350 Cep and V1647 Ori show similar photometric and spectral features. These include: a large amplitude outburst continuing several years, random fluctuations in brightness with amplitudes of a few tenths of a magnitude and timescales of several days (\citealt{GarciaAlvarez+etal+2011}), reddening of the color indices with decreasing brightness, connection to reflection nebulae (\citealt{Miranda+etal+1994}), an emission of line spectrum during the maximum light and a variable P Cyg profile for the H$\alpha$ line. FUors and EXors have been classified in terms of their wide range of available photometric and spectral properties, but their outbursts are thought to have been caused by an enhanced accretion rate. According to \cite{Aspin+2011}, the viewing inclination angle of the star/disk system can play a significant role in the observed spectral features, and therefore for the classification of the object as FUor or EXor. It is possible that the two types of eruptive variables FUors and EXors could be much closer in nature. \newpage \normalem | 14 | 4 | 1404.5902 |
1404 | 1404.0650_arXiv.txt | Among Type Ia supernovae (SNe~Ia) exist a class of overluminous objects whose ejecta mass is inferred to be larger than the canonical Chandrasekhar mass. We present and discuss the UV/optical photometric light curves, colors, absolute magnitudes, and spectra of three candidate Super-Chandrasekhar mass SNe--2009dc, 2011aa, and 2012dn--observed with the {\sl Swift} Ultraviolet/Optical Telescope. The light curves are at the broad end for SNe Ia, with the light curves of SN~2011aa being amongst the broadest ever observed. We find all three to have very blue colors which may provide a means of excluding these overluminous SNe from cosmological analysis, though there is some overlap with the bluest of ``normal'' SNe Ia. All three are overluminous in their UV absolute magnitudes compared to normal and broad SNe Ia, but SNe 2011aa and 2012dn are not optically overluminous compared to normal SNe Ia. The integrated luminosity curves of SNe 2011aa and 2012dn in the UVOT range (1600-6000 \AA) are only half as bright as SN~2009dc, implying a smaller $^{56}$Ni yield. While not enough to strongly affect the bolometric flux, the early time mid-UV flux makes a significant contribution at early times. The strong spectral features in the mid-UV spectra of SNe 2009dc and 2012dn suggest a higher temperature and lower opacity to be the cause of the UV excess rather than a hot, smooth blackbody from shock interaction. Further work is needed to determine the ejecta and $^{56}$Ni masses of SNe 2011aa and 2012dn and fully explain their high UV luminosities. | Supernovae \label{intro}} Type Ia Supernovae (SNe Ia) are important cosmological probes that first revealed the accelerating expansion of the universe \citep{Riess_etal_1998,Perlmutter_etal_1999}. The cosmological results rely on the normal SNe Ia whose brightness correlates with their light curve shapes and colors \citep{Phillips_etal_1999,Riess_etal_1996_mlcs,Goldhaber_etal_2001}, allowing them to be used as standardizable candles. Observations of similar but peculiar objects are useful for understanding the nature of the progenitor systems and the physics of the explosion, particularly how they might differ between objects. It is also important to understand objects which may be found in cosmological samples but do not follow the relationships between the luminosity and the light curve shape. The similar peak luminosities of SNe Ia suggested explosions of similar mass and energy. The widely-held theory is that a SN Ia results from the thermonuclear disruption of a Carbon-Oxygen white dwarf (CO-WD) as it approaches the Chandrasekhar limit. This could be due to accretion from a non-degenerate companion (also called the single degenerate scenario; \citealp{Whelan_Iben_1973}) or the disruption of a WD companion (also called the double degenerate scenario; \citealp{Webbink_1984, Iben_Tutukov_1984}). The nature of an SN Ia progenitor as a C-O WD (and admittedly for a single case) has only recently been confirmed by very early time observations of SN~2011fe \citep{Nugent_etal_2011,Bloom_etal_2012}. The WD mass at explosion might not need approach the Chandrasekhar limit, as helium shell detonations can trigger a core detonation in sub-Chandrasekhar mass progenitors \citep{Woosley_Weaver_1994,Fink_etal_2010,Kromer_etal_2010,Woosley_Kasen_2011}. The nature of the companion remains unknown, and recent results suggest that SNe Ia may result from both single degenerate and double degnerate systems. Early observations of many SNe Ia do not show the interaction expected \citep{Kasen_2010} if the SN explosion were to interact with a red giant (RG) companion \citep{Hayden_etal_2010_shock,Bianco_etal_2011, Mo_etal_2011, Brown_etal_2012_shock}. X-ray limits also rule out red giants due to the lack of shock interaction \citep{Russell_Immler_2012}. Pre-explosion, multi-wavelength, and extremely early observations of SN~2011fe rule out a RG \citep{Nugent_etal_2011,Li_etal_2011,Horesh_etal_2012,Margutti_etal_2012} and even a main sequence (MS) companion \citep{Bloom_etal_2012,Brown_etal_2012_11fe} for that object. Searches for the leftover companion in SNR 0509-67.5 rule out a non-degenerate companion \citep{Schaefer_Pagnotta_2012}. On the other hand, high resolution spectroscopy of nearby SNe has found a preference for blue shifted sodium absorption in about 20-25\% of SNe Ia \citep{Sternberg_etal_2011,Foley_etal_2012_prog,Maguire_etal_2013} and even variable absorption \citep{Patat_etal_2007,Simon_etal_2009} suggestive of a local CSM wind from a non-degenerate companion. PTF11kx observations showed signatures of a recurrent nova progenitor in a single degenerate system \citep{Dilday_etal_2012}. Thus, multiple channels might be required to create the explosions classified as SNe Ia. The idea that the accreting progenitor explodes as it approaches the Chandrasekhar mass has been challenged by a class of SNe that appear spectroscopically similar to SNe Ia but are overluminous for their light curve shape. Detailed modeling of the light curves appears to require more than a Chandrasekhar mass of ejected material. SN~2003fg was the first discovered \citep{Howell_etal_2006} with SNe 2006gz \citep{Hicken_etal_2007}, 2007if \citep{Scalzo_etal_2010, Yuan_etal_2010} and 2009dc \citep{Yamanaka_etal_2009,Tanaka_etal_2010,Silverman_etal_2011,Taubenberger_etal_2011,Kamiya_etal_2012,Hachinger_etal_2012} showing similarities. \citet{Scalzo_etal_2012} discovered five additional, similar objects in SN Factory observations, though only one was conclusively above the Chandrasekhar limit. Association with this subclass is sometimes based on spectroscopic similarity to others of the class, to a high inferred luminosity, or to actually modeling the light curve and determining a high ejecta mass. Variations exist amongst candidates of his subclass, which is not surprising given our limited understanding of their origin and relationship to normal SNe Ia. \citet{Maeda_Iwamoto_2009} highlight the observational differences between SNe 2003fg and 2006gz, two probable super-Chandarasekhar mass candidates. The most common means of estimating the mass from SNe Ia comes from the application of ``Arnett's Law'' \citep{Arnett_1982,Branch_1992}. At maximum light the luminosity output is approximately equal to the instantaneous rate of energy release from radioactive decay. Thus the peak bolometric luminosity is proportional to the mass of $^{56}$Ni synthesized in the explosion. The $^{56}$Ni can also be estimated from the late light curve \citep{Silverman_etal_2011} or nebular spectra \citep{Mazzali_etal_1997}. The total mass can be estimated based on energetics using the observed luminosities and expansion velocities and assumptions on the density profile (e.g. \citealp{Scalzo_etal_2012}). The mass can also be estimated by constructing models of various masses and explosion scenarios and comparing to the observed light curves \citep{Kamiya_etal_2012} and spectra \citep{Mazzali_etal_1997,Hachinger_etal_2012}. Not all of the luminosity necessarily comes from radioactive decay. Excess luminosity could also come from circumstellar interaction \citep{Taubenberger_etal_2013} or result from asymmetric explosions viewed at a favorable angle \citep{Hillebrandt_etal_2007}. Asymmetric explosions cannot explain the brightest of SC SNe, and spectropolarimetry of SN~2009dc implies no large scale asymmetries in the plane of the sky \citealp{Tanaka_etal_2010}). \citet{Maeda_etal_2009} find that the late time observations of SN~2006gz require less radioactive Ni than suggested from peak optical observations, drawing into question the overluminous nature of the event. They suggest that the luminosity is overestimated due to an over-correction for extinction. SC SNe are hot, high-energy explosions, so ultraviolet (UV) coverage is important to better measure the total luminosity and determine its origin, in particular whether it originates from shocks or simply a hot photosphere. The Ultraviolet/Optical Telescope (UVOT; \citealp{Roming_etal_2005}) on the Swift satellite \citep{Gehrels_etal_2004} presents an excellent opportunity to obtain unique, early-time UV data. This paper will focus on three objects: SN~2009dc--a well-studied member of the Super-Chandrasekhar mass SN class--and SNe 2011aa and 2012dn which share some characteristics. We will refer to these candidate super-Chandrasekhar SNe Ia as SC SNe below, though a firm mass determination will require more data and is beyond the scope of this work. Comparisons will focus on the differences and similarities between SN~2009dc and the less studied SNe 2011aa and 2012dn, and the differences of these three SC SNe compared to other SNe Ia. In Section \ref{obs} we discuss these three SC SNe and present UV/optical photometry and spectra from UVOT. In Section \ref{results} we compare the colors, absolute magnitudes, spectra, and integrated luminosities, comparing SNe 2011aa and 2012dn to 2009dc and the three to a larger sample of ``normal'' SNe Ia. In Section \ref{discussion} we discuss the results and summarize. | \label{discussion} One suggestion for the increased luminosity in SC SNe is shock interaction \citep{Fryer_etal_2010, Blinnikov_Sorokina_2010, Taubenberger_etal_2011, Hachinger_etal_2012}. \citet{Scalzo_etal_2012} suggest UV observations as a means to probe the influence of shock interactions on the early luminosity. \citet{Fryer_etal_2010} performed numerical calculations of the spectra and simulated UVOT light curves resulting from a double-degnerate SN Ia exploding within a shell of unaccreted material. While our candidate SC SNe Ia have peak luminosities comparable to those studied by \citet{Fryer_etal_2010}, the light curves shapes are much different. The light curve shapes can vary based on the amount and spatial distribution of the surrounding material, but the smoothness of the UV light curves and their qualitative similarity to the optical light curves suggest a photospheric origin. A photospheric origin for the emission is supported by the UV spectra of SNe 2009dc and 2012dn, which show stronger features in the MUV (below 2700 \AA) than seen in normal SNe Ia, while the flux from a hot shock would be relatively smooth and would dilute the photospheric features \citep{Hamuy_etal_2003}. On the other hand, the optical features are also much stronger than for SN~2007if, for which the top lighting of a shock was invoked as one explanation for its diluted features and high luminosity \citep{Scalzo_etal_2010}. The UV spectra of SNe 2009dc and 2012dn do not allow a smooth blackbody source for the excess flux. Such a spectrum might be expected from a high temperature shock with a hydrogen-rich circumstellar medium, as was used to explain the diluted features of SN~2002ic (\citealp{Hamuy_etal_2003}, see also \citealp{Branch_etal_2000}). A structured spectrum with emission and absorption, due to reprocessing of the shock emission or originating from a different composition, cannot be excluded. \citet{Hachinger_etal_2012} found adding a spectrum of the Ibn SN~2006jc to their theoretical SN Ia spectrum gave reasonable matches to the observed spectra of SN~2009dc. Higher quality UV spectra of Ibn and SC SNe are needed to perform similar tests in the UV. Nevertheless, photometric observations may already contain enough information to further constrain photometric (e.g. \citealp{Kamiya_etal_2012}) or spectroscopic modeling (e.g. \citealp{Hachinger_etal_2012}). While the optical light curves of SNe 2011aa and 2012dn are not dissimilar to normal SNe Ia (though extremely broad in the case of SN~2011aa), the NUV-optical and especially the MUV-NUV (or MUV-optical) colors are markedly different. The rest-frame UV also peaks earlier for SC than normal SNe Ia. Early rest-frame UV photometry might allow optically overluminous SNe such as SN~2009dc to be excluded from cosmological analysis. \citet{Scalzo_etal_2012} estimate the rate of SC SNe to be a few percent of all SNe Ia locally, but a bias could result from an evolutionary shift \citep{Taubenberger_etal_2011} if these are more common in the early universe than they are locally. \citet{Milne_etal_2013_z} show that the relative fractions of NUV-blue and NUV-red normal SNe Ia change with redshift. The origin of the UV diversity amongst normal and SC candidate SNe Ia may % point to ways to reduce the dispersion at longer wavelengths and understand potential biases in SN Ia standardization at different epochs in the history of the universe. Bolometric light curve comparisons between models and observations serve as an important diagnostic of allowed models and parameters. The creation of bolometric light curves, however, especially the treatment of missing wavelength ranges, varies greatly. Sometimes the NUV (or at least the ground based U band) is included, and the MUV may or may not be included. Often the UV portion of the flux is considered to be negligible (a reasonable assumption in some cases). If it is included, it is often set at a constant percentage of the flux. As shown here, there is also a lot of variation in the NUV and MUV flux fractions between various SNe Ia, and the fractions evolve quite significantly with time. The data given here will allow the bolometric light curves of these objects to be more accurately determined. For example, the falling UV fraction means that inclusion of the UV flux will broaden the pre-maximum rise of the bolometric flux. This could lead to a longer implied rise time if fit with a light curve template. This longer rise time may not be accurate, however, if the stretched light curve template did not include the UV in its construction. \citet{Kamiya_etal_2012} use multi-wavelength modeling to show the difference between a BVRI, UV-Optical-IR (UVOIR), and true bolometric light curve. The distinction between these is important. UV data will allow more constraints on the modelling. While we have pushed the knowledge of the UV behavior for SC SNe Ia $\sim8$ days earlier, the very earliest epochs would also be important for looking for the effects of shock interaction with a non-degenerate companion \citep{Kasen_2010,Brown_etal_2012_shock} or differences in the UV-optical flux evolution at the earliest times \citep{Brown_etal_2012_11fe}. As the UV-optical colors are still bluest at the first epochs observed, the bolometric contribution before then may be larger still and are in any case uncertain. Higher quality UV spectra at the earliest possible epochs will better probe the mechanism responsible for the excess UV emission and how to account for it in mass determinations. In summary, we have presented UV/optical photometry and spectroscopy for three SNe Ia, 2009dc, 2011aa and 2012dn, which have been suggested as candidate super-Chandrasekhar mass SNe Ia. While their optical properties are not dissimilar to normal SNe Ia, they are significantly bluer and more luminous in the UV than normal SNe Ia, with MUV luminosities about a factor of $\sim10$ higher. UV spectra of SNe~2009dc and 2012dn feature structure not expected for shock interaction, suggesting a photospheric origin of the excess UV luminosity. The UV is shown to contribute significantly (but still smaller than the optical) to the bolometric luminosity, especially at early times. The integrated luminosities of SNe 2011aa and 2012dn are much lower than 2009dc, however. This suggests a larger diversity in the class, if they are indeed in the same class, when considering UV and optical photometric and spectroscopic characteristics. A more detailed study of these SNe is required to determine if they were above the Chandrasekhar mass. | 14 | 4 | 1404.0650 |
1404 | 1404.3842_arXiv.txt | We report on the characterization of a number of AGB candidate stars identified with objective-prism plates of the Byurakan Observatory. Digitized photographic sky survey plates and recent CCD photometry have been used to improve the selection and distinguish variable and non-variable stars. Some comparisons among published catalog magnitudes are also made. Slit spectroscopy from the Asiago and Loiano Observatories allowed a firm spectral classification, separating C-Type, N-Type and normal M giants. Color-color plots using WISE, AKARI and 2MASS J-band data allow an efficient discrimination of spectral types, which can be used for the definition of larger statistical samples. \noindent \textbf{Keywords}: surveys; photographic plates; late type stars. | The large Schmidt telescope of the Byurakan Observatory (102/132/213 cm) with the 1.5 degree prism and 103aF emulsion made a successful survey of the northern sky in the 1970's. On these plates late type stars are easily recognized by their strong, nearly point-like, red emission while the blue part of the spectrum is very faint or even absent. This allows to build sample of AGB stars candidates, and/or to check the nature of infrared (IR) objects detected by space-based IR instruments like WISE, IRAS, AKARI or of ground-based surveys like 2MASS. The first Byurakan Survey plates and automatically extracted spectra are freely accessible at http://byurakan.phys.uniroma1.it/index.php (Mickaelian et. al. 2007). The AGB stars, due to their large luminosity, can be traced up to the limit of our Milky Way halo and therefore serve as tracers of its gravitational field and of dark matter distribution (see e.g. Mauron et al. 2013 and references therein). In this paper we report on an on-going survey program of AGB candidates in the northern hemisphere, aimed at obtaining a better spectral classification and exploring the possibilities of a classification from recent photometric infrared data from satellite surveys. Candidate stars were taken from the "Revised and Updated Catalogue of FBS Late-Type Stars"( Gigoyan and Mickaelian, 2012). | Photographic archives can still contribute to high quality science, provided that they can be easily accessed from the web. A fast consultation of images, even if of not high quality, may indeed help the researcher to have a quick response and decide if further investigation is worth to be made. The technology of plate scanners is rapidly evolving, as well as the availability of very massive electronic storage at low price, so the task of digitization is easier than just 10 years ago. A good storage of the original photographic material is therefore important, because in few years better scanners may be available allowing more accurate data retrieval. Putting on line at least the logbooks of the plate archives is mandatory making the old photographic data accessible and efficiently used. \vskip 3cm {\bf Acknowledgements} This publication makes use of data products from the Wide-field Infrared Survey Explorer, which is a joint project of the University of California, Los Angeles, and the Jet Propulsion Laboratory/California Institute of Technology, funded by the National Aeronautics and Space Administration. We thank the Directorate of the Loiano and Asiago Observatories for telescope time allocation | 14 | 4 | 1404.3842 |
1404 | 1404.6526_arXiv.txt | {The Chamaeleon molecular cloud complex is one of the nearest star-forming sites encompassing three molecular clouds (Cha~I, II, and III) with a different star-formation history, from quiescent (Cha~III) to actively forming stars (Cha~II), and reaching the end of star-formation (Cha~I).} {We aim at characterising the large-scale structure of the three sub-regions of the Chamaeleon molecular cloud complex by analysing new far-infrared images taken with the \emph{Herschel} Space Observatory.} {We derived column density and temperature maps using PACS and SPIRE observations from the \emph{Herschel} Gould Belt Survey, and applied several tools, such as filament tracing, power-spectra, $\Delta$-variance, and probability distribution functions of column density (PDFs), to derive physical properties.} {The column density maps reveal a different morphological appearance for the three clouds, with a ridge-like structure for Cha~I, a clump-dominated regime for Cha~II, and an intricate filamentary network for Cha~III. The filament width is measured to be around 0.12$\pm$0.04 pc in the three clouds, and the filaments found to be gravitationally unstable in Cha~I and II, but mostly subcritical in Cha~III. Faint filaments (\emph{striations}) are prominent in Cha~I showing a preferred alignment with the large-scale magnetic field. The PDFs of all regions show a lognormal distribution at low column densities. For higher densities, the PDF of Cha~I shows a turnover indicative of an extended higher density component, culminating with a power-law tail. Cha~II shows a power-law tail with a slope characteristic of gravity. The PDF of Cha~III can be best fit by a single lognormal.} {The turbulence properties of the three regions are found to be similar, pointing towards a scenario where the clouds are impacted by large-scale processes. The magnetic field could possibly play an important role for the star-formation efficiency in the Chamaeleon clouds if proven that it can effectively channel material on Cha~I, and possibly Cha~II, but probably less efficiently on the quiescent Cha~III cloud.} | \label{intro} \begin{figure*}[ht] \begin{center} \hspace{0.0cm}\includegraphics[angle=0,width=18cm]{fig1.pdf} \caption[]{SPIRE 250~$\mu$m map with extinction contours \citep{Schneider2011} overlaid from 2 to 10~mag, in steps of 2~mag. The arrows show the general orientation of the magnetic field as measured in Cha~I by \citet{Whittet1994} and \citet{McGregor1994}. Note the preferential alignment of the faint \emph{striations} in Cha~I, less evident in Cha~II or Cha~III where the surrounding diffuse emission takes a variety of orientations. In the region eastwards of Cha~III, no \emph{striations} are detected in the \emph{Herschel} maps. } \end{center} \label{spire} \end{figure*} Molecular clouds in the Galaxy bear a complex spatial structure highlighted by the unprecedented sensitivity and resolution of the \emph{Herschel} Space Observatory \citep{Pilbratt2010} imaging observations, mainly from large survey programs \citep{Andre2010,Motte2010,Molinari2010}. The recent advances in the census of starless and prestellar cores makes it possible to study their connection to the large scale structure of molecular clouds. Observations show that the prestellar cores identified with \emph{Herschel} are preferentially found within gravitationally unstable filaments \citep[e.g.,][K\"onyves et al., \emph{in prep.}]{Andre2010,Andre2014,Polychroni2013,Konyves2010}, while massive protostellar dense cores and star clusters tend to be found at the junctions of dense filaments \citep[e.g.,][]{Hennemann2012,Schneider2012,Peretto2013}. The observed orientation of filaments suggests that they are aligned with the magnetic field \citep[e.g., in the massive star-forming region DR21,][]{Schneider2010}. Recently, \citet{Palmeirim2013} discovered that low-density filaments (\emph{striations}) in the Taurus region are also preferentially oriented along the magnetic field. These findings suggest that the initial conditions that favour star formation are closely linked to the spatial structure of a molecular cloud. Identifying the processes responsible for the fragmentation of clouds into dense filaments and their subsequent evolution, is therefore paramount to the understanding of core formation, and ultimately, the dependence of core and stellar masses on the large-scale properties of the interstellar medium. The Chamaeleon molecular cloud complex is one of the nearest star-forming sites located at a distance of 150$-$180~pc \citep{Whittet1997}. It contains the Cha I, II, and III (Fig.~\ref{spire}), as well as the Musca clouds (Cox et al., \emph{in prep.}). \citet{Mizuno2001} showed in their $^{12}$CO~(1$\to$0) survey that the complex is spatially and kinematically coherent with emission in the range of $-$4 to 6~km~s$^{-1}$. Cha~I is the most active star-forming region with a young stellar population of over $\sim$200 members \citep{Luhman2008,Winston2012} with a median age of $\sim$2~Myr. Cha~II ($\sim$4~Myr) has a smaller population of $\sim$60 young stellar objects (YSOs) \citep{Spezzi2008,Spezzi2013}, and no YSO has been found in Cha~III \citep{Belloche2011a}. Using the Large APEX Bolometer Camera (LABOCA), \citet{Belloche2011a,Belloche2011b} mapped the dust continuum emission at 870~$\mu$m in Cha~I and III to single out the possible causes of such different star-formation activity. In Cha~I, the low number of candidate prestellar cores and protostars, as well as the high global star formation efficiency, were interpreted as signs that star formation might be at its end in this cluster, whereas in Cha~III evidence for the on-set of star formation was found. The Chamaeleon dark clouds were observed with \emph{Herschel} as part of the Gould Belt survey \citep[hereafter, HGBS,][]{Andre2010}. These observations are presented in Sect.~\ref{observations}, and represent an homogeneous dataset in the far-IR across a large area of the cloud complex that allows us to characterize the large-scale structure and extended emission of the three clouds, to better understand their different star formation history. In Sect.~\ref{methods}, we present the different analysis tools employed, such as filament tracing, power spectra, $\Delta$-variance, and probability distribution functions of the column density (PDFs). The properties derived, such as filamentary \emph{vs.} clumpy structure, level of turbulence and energy injection, density structure, are presented in Sect.~\ref{results} and discussed in Sect.~\ref{discussion}, with the conclusions presented in Sect.~\ref{conclusion}. \begin{figure*}[ht] \centering \includegraphics[height=6.cm]{fig2a.pdf} \includegraphics[height=6.cm]{fig2b.pdf} \includegraphics[height=6.cm]{fig2c.pdf} \includegraphics[height=6.cm]{fig2d.pdf} \includegraphics[height=6.cm]{fig2e.pdf} \includegraphics[height=6.cm]{fig2f.pdf} \caption{Column density (N$_{H_2}$(cm$^{-2})$) \emph{(top)} and temperature (K) \emph{(bottom)} maps of Cha~I, II, and III \emph{(from left to right)} with the skeletons tracing the most prominent filamentary structure superimposed in black.} \label{col} \end{figure*} | \label{conclusion} New \emph{Herschel} photometric observations of the three Chamaeleon clouds (I, II, and III) taken with the PACS and SPIRE instruments in an homogeneous way were presented in this paper. They were analysed with a set of tools to characterise quantitatively the large scale structure and extended dust emission, and study their possible relation to the accentuated differences seen today amongst the three clouds, with Cha~I likely at the end of star-formation, Cha~II actively forming stars, and Cha~III in a quiescent state. The column density and temperature maps derived from the \emph{Herschel} data reveal important morphological differences for the three clouds, with a ridge-like structure for Cha~I surrounded by faint filaments (\emph{striations}) aligned with the large-scale magnetic field, a clump-dominated regime for Cha~II, and a complex low-density filamentary network for Cha~III. Filamentary-like structures share a common width ($\sim$0.12$\pm$0.04~pc) consistent with values inferred from observations of other star-forming regions \citep[e.g.,][]{Arzoumanian2011,Andre2014}. However, only in Cha~I and II filaments are found to be predominately gravitationally unstable. All regions show a PDF described by a lognormal distribution for low column densities with a width of $\sim$0.45 to 0.6 mag and a peak at A$_{V}$$\sim$2~mag. For higher column densities, the PDFs show significant differences, with Cha~II being the only region where a classical single power-law tail with a slope indicative of free-fall collapse is seen. We compared the PDFs to the results from hydrodynamic simulations of \citet{Federrath2012}, and conclude that they are broadly described by models where $\mathcal{M}$=3, 5, or 10 with solenoidal or mixed forcing (with or without magnetic field) and low SFE. Overall, the turbulence properties of the three regions do not show large differences, lending strength to a scenario where the clouds are impacted by common large-scale processes. We emphasise, however, that an alignment of faint filaments peripheral to dense structures with the magnetic field is clearly seen in Cha~I. Similar preferential distributions have been found in other star-forming regions \citep[e.g.,][]{Palmeirim2013}. Future results from the \emph{Planck} mission on polarisation will quantify the relation between the magnetic filed and the structures in Cha~II and Cha~III. If proven that \emph{striations} are important channels of accretion of ambient material into filaments, the magnetic field could play an important role in shaping the differences in the star-formation history of the Chamaeleon complex regions. | 14 | 4 | 1404.6526 |
1404 | 1404.5713_arXiv.txt | We demonstrate how the metallicities of red supergiant (RSG) stars can be measured from quantitative spectroscopy down to resolutions of $\approx$ 3000 in the \jband. We have obtained high resolution spectra on a sample of the RSG population of \ob, a double cluster in the solar neighborhood. We show that careful application of the {\sc marcs} model atmospheres returns measurements of \met\ consistent with solar metallicity. Using two grids of synthetic spectra$-$one in pure \lte\ and one with \nlte\ calculations for the most important diagnostic lines$-$we measure \met\ = \ltezfit\ (\lte) and \met\ = \nltezfit\ (\nlte) for the sample of eleven RSGs in the cluster. We degrade the spectral resolution of our observations and find that those values remain consistent down to resolutions of less than $\lambda/\delta\lambda$ of 3000. Using measurements of effective temperatures we compare our results with stellar evolution theory and find good agreement. We construct a synthetic cluster spectrum and find that analyzing this composite spectrum with single-star RSG models returns an accurate metallicity. We conclude that the RSGs make ideal targets in the near infrared for measuring the metallicities of star forming galaxies out to 7-10 Mpc and up to ten times farther by observing the integrated light of unresolved super star clusters. | Measuring metallicities in star-forming galaxies is a ubiquitous goal across the field of extragalactic astronomy. The evolutionary state of a galaxy is imprinted in the central metallicity and radial abundance gradient of iron- and $\alpha$-group elements. Observed trends in these measurements across ranges of galactic mass, redshift, and environment constrain the theory of galaxy formation and chemical evolution. Central metallicity is dictated by galactic mass, a relationship encoded by the initial properties and evolution of these objects \citep{1979A&A....80..155L,2004ApJ...613..898T,2008A&A...488..463M}. Radial metallicity gradients provide a wealth of information needed to describe the complex dynamics of galaxy evolution including clustering, merging, infall, galactic winds, star formation history, and initial mass function \citep{2000MNRAS.313..338P,2004cmpe.conf..171G,2008A&A...483..401C,2009A&A...505..497Y,2009MNRAS.398..591S,2004MNRAS.349.1101D,2007MNRAS.374..323D,2008MNRAS.385.2181F,2007ApJ...655L..17B,2007MNRAS.375..673K,2009MNRAS.399..574W}. The pursuit of these scientific goals has been undermined by the difficulty of obtaining reliable metallicities. Investigations tend to rely on spectroscopy of the emission lines of \hii\ regions. These methods require empirical calibration and choosing different commonly used calibrations yields varying and sometimes conflicting results from the same set of observations. Both the slope and absolute scaling of metallicity are susceptible to choice of calibration: the mass-metallicity gradient across all galaxies and the radial gradients within individual galaxies can change from steep to flat while the overall metallicity can shift by a factor of up to four \citep{2008ApJ...681.1183K,2008ApJ...681..269K,2009ApJ...700..309B}. Even the more physical ``T$_e-$based method'' (which utilizes auroral lines to remove the need for ``strong line'' calibrations) is potentially subject to biases$-$especially in the metal rich regime characteristic of the disks of all massive spiral galaxies \citep{2014arXiv1401.4437B,2005A&A...434..507S,2005A&A...441..981B,2010MNRAS.401.1375E,2012MNRAS.427.1463Z}. One technique which avoids the uncertain calibrations of the ``strong line'' \hii\ region method is the quantitative spectroscopy of supergiant stars. Blue supergiants have become a powerful tool for measuring metallicities, gradients, and distances to galaxies in and beyond the Local Group (WLM \textendash{} \citealt{2006ApJ...648.1007B,2008ApJ...684..118U}; NGC 3109 \textendash{} \citealt{2007ApJ...659.1198E}; IC1613 \textendash{} \citealt{2007ApJ...671.2028B}; M33 \textendash{} \citealt{2009ApJ...704.1120U}; M81 \textendash{} \citealt{2012ApJ...747...15K}). This technique, while extremely promising, may also be subject to systematic uncertainties and needs to be checked by independent methods. Moreover, it requires optical spectroscopy. However, next generation telescopes such as the TMT and E\textendash{}ELT will be optimized for observations at infrared wavelengths, using adaptive optics supported multi object spectrographs. Thus, we need bright abundance tracers which radiate strongly in the IR. Such stars--including red giants, the asymptotic giant branch, and red supergiants--will have a clear advantage in the future. \begin{figure*} \begin{centering} \includegraphics[width=6in]{f1.eps} \caption{Spectral library of RSGs observed at high resolution with IRCS on Subaru. The main diagnostic atomic lines are labeled. Best fitting \nlte\ models are over plotted in red. The Mg\,{\sc i} line is not included in the fit because it is calculated in \lte\ but subject to strong \nlte\ effects. \nlte\ calculations for Mg\,{\sc i} will be implemented soon. Plots are arranged by spectral type (see Table~\ref{tbl:perOB1}).} \label{fig:speclib} \end{centering} \end{figure*} \begin{figure*} \begin{centering} \includegraphics[width=6in]{f2.eps} \caption{Spectral library from Figure~\ref{fig:speclib} downgraded to a resolution of R=3000. The main diagnostic atomic lines are labeled. Best fitting \nlte\ models are over plotted in red. The Mg\,{\sc i} line is not included in the fit because it is calculated in \lte\ but subject to strong \nlte\ effects. \nlte\ calculations for Mg\,{\sc i} will be implemented soon. Plots are arranged by spectral type (see Table~\ref{tbl:perOB1}).} \label{fig:speclib2} \end{centering} \end{figure*} \begin{figure} \begin{centering} \includegraphics[width=8.5cm]{f3.eps} \caption{The effects of \nlte\ corrections on diagnostic lines. Left panel shows a model at \met\ = $-$0.25, right panel \met\ = $+$0.25} \label{fig:lvsnl} \end{centering} \end{figure} \begin{figure} \begin{centering} \includegraphics[width=8.5cm]{f4a.eps} % \includegraphics[width=8.5cm]{f4b.eps} \caption{Plot of continuum region between two strong atomic features at spectral resolution of 3000. Each model has \teff=4000 \logg=0.0, and $\xi$ = 4.0. Top panel: models at four values of \met\ to demonstrate the \pc. Red squares mark the continuum points used for each model. Bottom panel: each model is scaled to that of \met=0.0 assuming that it resembles the data set. The variable depth of atomic spectral features as a function of metallicity is still clearly seen. In addition, weak line features strengthen with metallicity and provide additional information with increasing metallicity.} \label{fig:pc} \end{centering} \end{figure} The extremely luminous red supergiant stars (RSGs)\textemdash{}which emit 10$^5$ to $\sim$10$^6$ L/L$_\odot$ largely in the infrared \citep{Humphreys:1979p3252}\textemdash{}thus become ideal targets for measuring extragalactic cosmic abundances. Complications due to the densely packed spectral features synonymous with the cool, extended atmospheres of RSGs are minimized in the \jband. Here the dominant features are isolated atomic lines of iron, titanium, silicon, and magnesium. Molecular lines of OH, H$_2$O, CN, and CO manifest weakly or not at all in this bandpass. A new technique proposed by \cite{2010MNRAS.407.1203D} (henceforth DFK10) has demonstrated that quantitative, medium resolution spectroscopy (R [$\lambda/\delta\lambda$] $\sim$2000) in the \jband\ can determine metallicities accurate to $\sim$0.15 dex for a single RSG. While a principal limitation of the quantitative spectroscopy of stars is distance, these supergiant studies using 8-meter class telescopes have the potential to be extended to $\sim$10 Mpc \citep{2011A&A...527A..50E}. \begin{figure*} \begin{centering} \includegraphics[width=6.5in]{f5.eps} \caption{2D contour plots used to extract fit parameters in the analysis procedure. The smooth color gradient is an interpolated $\chi^2$ map, with dark representing lower values (better fits). White contour lines depict fit areas of 1, 2, and 3 $\sigma$ as determined by monte carlo sampling. The blue point at the intersection of blue lines shows the minimum $\chi^2$ in each 2D slice.} \label{fig:fitmethod} \end{centering} \end{figure*} The \jband\ technique is thus poised to study a substantial volume of the local universe, one containing groups and clusters of galaxies. The determination of accurate abundances for the RSG populations of star forming galaxies in this volume will provide an unparalleled observational constraint for models of galaxy formation and evolution. An increased utilization of supergiant stars may also aid in the proper development of the observationally efficient \hii-region methods while providing independent alternate measurement technique to the blue supergiants. Still, DFK10 is a pilot study of the \jband\ technique and the analysis methods to best study these stars requires careful development and testing. Studies of RSGs have classically required high resolutions (R$\sim$20,000) in the H\textendash{}band in order to separate and study the dense forest of atomic and molecular features present throughout their spectra. Part of this requirement is driven by the scientific desire to study stellar evolution, for which abundances of C, N, and O are important. The \jband\ technique returns no information specific to CNO processing and in exchange avoids the high observational overloads inherent to such studies. This repurposing for extracting global chemical enrichment at modest resolution is novel. Multiple facets of ongoing research investigate the limitations and systematic uncertainties of the technique in great detail. \cite{2013ApJ...767....3D} provide a thorough investigation of the temperature scale of RSGs in the LMC and SMC and conclude that previous work at optical wavelengths measure effective temperatures which are too cool for these RSGs. % They find that {\sc marcs} models which fit the strong optical TiO bands produce too little flux in the infrared to fit observed RSG spectral energy distributions. This discrepancy manifests in low measurements of effective temperature when fitting is performed with optical spectroscopy alone. This problem greatly reduced in the near-IR which correspond to deeper atmospheric layers. Additional research is assessing the significance$-$and observational effects$-$of the local thermodynamic equilibrium (\lte) calculations for synthetic spectra produced from the {\sc marcs} models. Departures from \lte\ have been calculated for iron and titanium \citep{2012ApJ...751..156B} and silicon lines \citep{2013ApJ...764..115B} in the \jband. Due to the low density environments in the extended atmospheres of RSGs, \nlte\ effects are noticeable and can be significant. For this work we have access to synthetic spectra calculated in both \lte\ ({\sc turbospectrum} $-$ \citealt{1998A&A...330.1109A,2012ascl.soft05004P}) and with iron, titanium, and silicon lines in \nlte\ using the results from \cite{2012ApJ...751..156B,2013ApJ...764..115B}. The aim of this paper is to carefully study the proposed methods of DFK10 and develop a proper understanding of the strengths, limitations, and systematics of the technique. The ideal target for such a study is a nearby coeval population of RSGs in the Galaxy, such that we may study the stars as individual objects and test the potential of utilizing distant super star clusters (SSCs) in which the stellar population becomes an unresolved point source. Theoretical predictions by \cite{2013MNRAS.430L..35G} show that in young SSCs the RSG population dominates the near-infrared flux. In this case the metallicity of the cluster could be extracted by studying the entire cluster as a single RSG. In order to accomplish these goals we target a galactic population of RSGs in the \ob\ double cluster (henceforth \obs) by performing quantitative spectroscopy on high resolution, high precision spectra collected using the Subaru Telescope atop Mauna Kea. The presence of a large population of supergiant stars limits the age of \obs\ to tens of millions of years, and offers a laboratory for the full range of stellar astrophysics--from IMF to post-main sequence stellar evolution. \cite{2010ApJS..186..191C} present a careful photometric and spectroscopic study of the double cluster and refine the physical parameters of this system. They find an age of 14$\pm$1 Myr and estimate a minimum total stellar mass of 20,000 M$_\odot$. Ages are determined using three methods which return results in good agreement: main sequence turnoff fitting, the luminosities of red supergiants in the clusters, and pre main sequence isochrone fitting. Solar metallicity is a sensible assumption for such a young population in the Milky Way, and studies of the B and A population of supergiant and giant stars$-$while incomplete$-$find solar or slightly sub-solar abundances. Our high resolution spectra of eleven RSGs in \obs\ provide an ideal dataset for testing multiple aspects of this project. This paper is organized as follows. In \S\ref{sec:obs} we discuss the observation and reduction of our spectral database. \S\ref{sec:technique} contains a description of our atmosphere models and synthetic spectra as well as an outline of the analysis method we have developed. We discuss the results of our fitting in \S\ref{sec:results} We discuss and summarize the results of this work in \S\ref{sec:discussion}. \begin{deluxetable*}{lcccccl} \tablewidth{0pt} \tablecaption{Perseus OB-1 Red Supergiants} \tablehead{ \colhead{Target} & \colhead{RA} & \colhead{DEC} & \colhead{m$_{\rm{V}}$} & \colhead{m$_{\rm{J}}$} & \colhead{m$_{\rm{H}}$} & \colhead{SpT}} \startdata BD+59 372 & 01 59 39.66 & +60 15 01.9 & 9.30 & 5.33 & 4.20 & K5-M0 I \tablenotemark{a} \\ % BD+56 595 & 02 23 11.03 & +57 11 58.3 & 8.18 & 4.13 & 3.22 & M1 I \tablenotemark{a} \\ % HD 14404 & 02 21 42.41 & +57 51 46.1 & 7.84 & 3.56 & 2.68 & M1Iab \tablenotemark{b} \\ % HD 14826 & 02 25 21.86 & +57 26 14.1 & 8.24 & 3.47 & 2.47 & M2 I \tablenotemark{a} \\ HD 236979 & 02 38 25.42 & +57 02 46.2 & 8.10 & 3.26 & 2.30 & M2 I \tablenotemark{a} \\ HD 13136 & 02 10 15.79 & +56 33 32.7 & 7.75 & 3.00 & 2.14 & M2 Iab \tablenotemark{b} \\ HD 14270 & 02 20 29.00 & +56 59 35.2 & 7.80 & 3.38 & 2.48 & M2.5 Iab \tablenotemark{b} \\ % BD+56 724 & 02 50 37.89 & +56 59 00.3 & 8.70 & 3.10 & 2.00 & M3 Iab \tablenotemark{b} \\ % HD 14469 & 02 22 06.89 & +56 36 14.9 & 7.55 & 2.82 & 1.93 & M3-4 I \tablenotemark{a} \\ % BD+56 512 & 02 18 53.28 & +57 25 16.8 & 9.20 & 3.68 & 2.68 & M3 I \tablenotemark{a} \\ % HD 14488 & 02 22 24.30 & +57 06 34.4 & 8.50 & 3.05 & 2.11 & M4 I \tablenotemark{a} \enddata \tablecomments{Target list for calibration of low resolution \jband\ RSG metallicity extraction. m$_V$ values are adopted from \citealt{1992A&AS...94..211G}, m$_J$ and m$_H$ from 2MASS \citep{2006AJ....131.1163S}.} \tablenotetext{a}{\,Spectral type from \cite{2005ApJ...628..973L}.} \tablenotetext{b}{\,Spectral type from \cite{1992A&AS...94..211G}.} \label{tbl:perOB1} \end{deluxetable*} | In this paper we have tested the \jband\ technique for extracting metallicity information from modest resolution spectra of RSGs. Through a careful suite of tests we have demonstrated the precision and accuracy of the technique. We obtain reliable abundances in agreement with high resolution, high signal to noise spectroscopy of young massive B-stars in the solar neighborhood. Using the advantage that all of our RSGs formed within a stellar cluster we test our derived parameters against predictions of stellar evolution theory for a cluster of mass and age of \obs. Our results are in good agreement with such theoretical work. We thus confirm the technique presented in DFK10 and show that it remains stable down to resolutions of R$\approx$3000. This provides a reliable method to determine extragalactic metallicities from individual RSGs to distances of 7-10 Mpc with existing telescopes and instruments. Both Keck (MOSFIRE) and the VLT (KMOS) have multi object spectrographs in the near-IR which operate above resolutions of 3000. With these instruments and the \jband\ technique, RSGs across the entire disks of star forming galaxies can be observed efficiently. By utilizing the large populations of RSGs in young, spatially unresolved SSCs we can extend the applicability of the \jband\ technique out to distances ten times greater with the same instruments. Thus SSCs may allow us to reach beyond the local group and measure the metallicities of star forming galaxies from the stars themselves instead of relying on existing techniques which are empirically calibrated. We note that low resolution work is now needed in targets expected to be sub solar and super solar in metallicity. The successful application of the \jband\ technique in such cases would pose the methods tested in this paper to study the metallicity evolution of star forming galaxies in a large volume of the nearby universe. | 14 | 4 | 1404.5713 |
1404 | 1404.2902_arXiv.txt | {The cataclysmic variable V893~Sco is an eclipsing dwarf nova which, apart from outbursts with comparatively low amplitudes, exhibits a particularly strong variability during quiescence on timescales of days to seconds.} {The present study aims to update the outdated orbital ephemerides published previously, to investigate deviations from linear ephemerides, and to characterize non-random brightness variations in a range of timescales.} {Light curves of V893~Sco were observed on 39 nights, spanning a total time base of about 14 years. They contain 114 eclipses which were used to significantly improve the precision of the orbital period and to study long-term variations of the time of revolution. Oscillations and similar brightness variations were studied with Fourier techniques in the individual light curves.} {The orbital period exhibits long-term variations with a cycle time of 10.2 years. They can be interpreted as a light travel time effect caused by the presence of a giant planet with approximately 9.5 Jupiter masses in a 4.5 AU orbit around V893~Sco. On some nights transient semi-periodic variations on timescales of several minutes can be seen which may be identified as quasi-periodic oscillations. However, it is difficult to distinguish whether they are caused by real physical mechanisms or if they are the effect of an accidental superposition of unrelated flickering flares. Simulations to investigate this question are presented.}{} | \label{Introduction} Cataclysmic variables (CVs) are well known to be short period interacting binary systems where a Roche-lobe filling star, the secondary, transfers matter via an accretion disk to a white dwarf primary. The structure of CVs can best be studied in systems where the secondary, which is faint and contributes in most cases only negligibly to the optical light, eclipses periodically the bright accretion disk and the white dwarf. In systems like these it is possible not only to derive the orbital period easily and accurately, but also to set more stringent limits on many fundamental system parameters. Moreover, eclipses frequently can be used as a tool to study structural details of the binary system and its components. \object{V893 Sco} is a member of the dwarf nova subclass of CVs. It was identified as a variable star and classified by Satyvoldiev (\cite{Satyvoldiev}) but got lost thereafter. Only much later, in 1998, Kato et al.\ (\cite{Kato}) re-identified V893~Sco. Soon thereafter, Bruch et al.\ (2000; hereafter referred to as Paper~I) published a photometric study, reporting the discovery of eclipses and deriving an orbital period of $1^{\rm h}\, 49^{\rm m}\, 23^{\rm s}$ which makes the system a member of short-period CVs located below the famous gap in the orbital period distribution of cataclysmic variables. Moreover, they found strong cycle-to-cycle variations concerning the mean magnitude, the strength of the orbital hump, the presence of an intermediate hump, and the amplitude and minimum depth of the eclipses. Time-resolved optical spectroscopic studies of V893~Sco were published by Matsumoto et al.\ (\cite{Matsumoto}) and Mason et al.\ (\cite{Mason}) and show that the system -- in spite of exhibiting some peculiarities -- is spectroscopically similar to many other CVs of the same kind. Thorstensen (\cite{Thorstensen}) measured trigonometrically a distance of $153^{+68}_{-35}\, {\rm pc}$. Mukai et al.\ (\cite{Mukai}) observed V893~Sco in x-rays with the {\em Suzaku} satellite and found partial eclipses. Warner et al.\ (\cite{Warner03}) report on the detection of quasi-periodic oscillations (QPOs) [which have also been seen by Bruch et al.\ (2000)] and dwarf nova oscillations (DNOs), and Pretorius et al.\ (\cite{Pretorius}) claim to have observed an instant with DNO activity. In continuation of the observations discussed in \cite{PaperI}, V893~Sco was observed regularly over the past 14 years. In this study I present some of the results of this long-term effort. In Sect.~\ref{Observations}, the reader is introduced to the observations. The photometric state during which they were taken and the characteristics of the long-term light curve are briefly discussed in Sect.~\ref{Photometric state}. The partially eclipsed x-ray source in V893~Sco (Mukai et al., \cite{Mukai}) is expected to be located in the immediate vicinity of the white dwarf (e.g.\ a boundary layer). Therefore, the suspicion of Bruch et al.\ (\cite{PaperI}), based on the strong variability of the amplitude and minimum depth of the eclipses, that the eclipsed body is the hot spot and that the centre of the accretion disk and the white dwarf remain uneclipsed, is not tenable. This is the topic of Sect.~\ref{Eclipse profile} which is dedicated to an analysis of the eclipse profile. Mukai et al.\ (\cite{Mukai}) already noticed that the ephemerides derived in \cite{PaperI} did not well predict the eclipse times at the epoch of their x-ray observations. Thus, an update is required. This is done in Sect.~\ref{Ephemerides} using data collected over a vastly longer time base. The new ephemerides reveal cyclic period variations which are investigated in detail and interpreted in Sect.~\ref{The cyclic period variations}. The presence of a third body with the mass of a giant planet can explain the observations. The detection of QPOs and DNOs by Bruch at al.\ (\cite{PaperI}), Warner et al.\ (\cite{Warner03}) and Pretorius et al.\ (\cite{Pretorius}) justifies a frequency analysis of the light curves which is performed in Sect.~\ref{Frequency analysis}. Here, I also address the question of the reliability of peaks in power spectra in the presence of strong flickering activity. Finally, the results of this study are summarized in Sect.~\ref{Conclusions}. | \label{Conclusions} I have presented results of a long-term photometric study of the eclipsing dwarf nova V893~Sco, focussing on a new determination of the eclipse ephemerides and their interpretation, and on possible oscillations in the brightness of the system. The main conclusions can be summarized as follows: \begin{enumerate} \item Normal outbursts with a comparatively low amplitude are frequent but the absence of superoutbursts during 13 years of monitoring is unusual for a dwarf nova with a period below the gap in the orbital period distribution of cataclysmic variables. The likelihood for superoutbursts to hide in the observational gaps is found to be small, unless the supercycle is close to one year (or multiples thereof) and all superoutburst occur around conjunction of V893~Sco with the sun. \item The eclipse profile is strongly variable in shape, amplitude and minimum magnitude. Large differences in the visibility of the bright spot in the profiles of individual eclipses show that it is outshone by the strong ubiquitous flickering. On average, about 36\% of the total light of V893~Sco is occulted during mid-eclipse. \item Updated orbital ephemerides are calculated from eclipse timings, leading to a significantly more precise value of the orbital period. Simple linear ephemerides cannot adequately describe the observations. While a quadratic term is not significant, the orbital period undergoes a cyclic variation which leads to a sinusoidal modulation of 22.3 sec in the $O-C$ diagram with a period of 10.2 years. \item These cyclic variations can be interpreted as a light travel time effect if the presence of a giant planet with a mass of the order of 9.5 Jupiter masses at a distance of 4.5 AU from the binary system is postulated. It was shown that a planet with these characteristics can survive the common envelope phase which must have preceded the appearance of V893~Sco as a CV if the mass of the white dwarf in the system is on the low side of the distribution of white dwarf masses in cataclysmic variables ($\approx 0.5 - 0.6\, M_{\sun}$). \item A search for oscillations in the light curves of V893~Sco was performed. At low frequencies, no persistent signal other than a modulation with the orbital period and its first harmonic was detected. At high frequencies a search of DNO-type oscillations met with no success. More interesting and also more difficult to interpret are oscillations at intermediate frequencies (timescale: $2^{\rm m} \le \tau \le 15^{\rm m}$). In some light curves transient oscillations at frequencies between 0.5 and 3 mHz are seen for part of the time which may be identified with QPOs. However, an accidental superposition of unrelated flickering flares may also lead to the observed signals. Simulations were performed in order to gain insight to which degree random fluctuation can mimic QPOs. To this end randomized version of real light curves were analysed. It was shown that in a non-negligible fraction of cases their power spectra contained signals which can be mistaken as being due to QPOs. The simple detection of an apparently significant signal in the power spectrum of a light curve containing strong flickering is therefore not sufficient to claim the presence of QPOs\footnote{The simultaneous presence of DNOs can help (Warner, \cite{Warner04}). Their period is normally about 15 times less than the period of the QPOs. If DNOs at two different frequencies are seen at the same time, their beat period corresponds in many cases to the QPO period.}. Even so, it is shown to be unlikely that all apparent oscillations seen in the present ensemble of light curves can be explained by a chance superposition of unrelated events. Therefore at least some of them must have a physical origin. The question to which degree QPOs and flickering are really conceptually different is raised. \end{enumerate} | 14 | 4 | 1404.2902 |
1404 | 1404.5655_arXiv.txt | We explore features of a 3D Fermi liquid near generalized Pomeranchuk instabilities using a tractable crossing symmetric equation method. We approach the instabilities from the ordered ferromagnetic phase. We find ``quantum multi-criticality" as approach to the ferromagnetic instability drives instability in other channel(s). It is found that a charge nematic instability precedes and is driven by Pomeranchuk instabilities in both the $\ell=0$ spin and density channels. | \label{sec:ded} \vspace{-0.5em} (\small{by {Khandker Quader)}\\ \noindent ``{\it We are like dwarfs on the shoulders of giants. \\ So that we can see more than they, \\ And things at a greater distance,\\ Not by virtue of any sharpness of sight on our part,\\ Or any physical distinction,\\ But because we are carried high, \\ And raised up by their giant size}" \\ \noindent \small{(Bernard of Chartres/John Salisbury (12th century); Isaac Newton (17th century))}\\[-1em] Much of the many-fermion physics that treats short-range underlying interaction and longer-range quantum fluctuations on the same footing are rooted in the bold, seminal ideas of Gerry that some of us, as his students, had the great fortune of learning first-hand from him. Over the years, Gerry's ``induced interaction" edifice gave us the confidence and guidance to build this into a tractable crossing-symmetric theory, and apply to interacting fermion problems in condensed matter and nuclear physics. This contribution is dedicated to the memory of Gerry! | \label{sec:summ} In the TCSE method, an arbitrary underlying interaction is treated on the same footing as quantum fluctuations. In our model, only $\ell=0$ fluctuations are taken into account. $F_{0}^{s,a}$ can then be thought of as a sum of the direct interaction and quantum fluctuations. Upon approach to PIs's (with increasing $U$), quantum fluctuations are enhanced in both strong and weak (multicritical) FM branches of solutions, as evident from discussions and figures above. In the spin channel, fluctuations work together to cancel most of the contribution from the driving interaction, which leads to a small $F_{0}^{a}$, where by ``small", we mean ``close to $F_{0}^{a}$ approaching the Pomeranchuk instability. In FM FL theory, ferromagnetism is related to $F_{0}^{a}$ as $m\sim[1+F_{0}^{a}]^{1/2}$. In the multicritical branch of solutions, this quantity becomes weaker for increasing direct interaction, as opposed to becoming stronger, as in the strongly FM branch. This counterintuitive behavior of $F_{0}^{a}$ is due to the large fluctuations which are opposite in sign to the direct interaction, leading to a small $F_{0}^{a}$, which gets even smaller for increasing $U$ (i.e. approach to PI) due to enhancement of fluctuations in this area. In the density channel, the density fluctuations, together with the driving interaction, compete against spin fluctuations to give a small $F_{0}^{s}$, where by ``small", we mean close to the $F_{0}^{s}$ PI. Thus, this enhancement, interplay, and feedback of quantum fluctuations results in multicritical behavior in which both channels simultaneously approach a GPI. We have found that with explicit $q$-dependence in $F_{0}^{a}$, which allows $F_{0}^{a}$ to approach its $q=0$ PI ($F_{0}^{a}=-1$), the fluctuations in the density channel do not change qualitatively. However, the spin fluctuations in the spin channel are stronger and compete more with density fluctuations. Fluctuations in spin and density channels also affect the higher $\ell$ FL parameters, as evidenced by our discussions on approach to nematic PI, effective mass and pairing amplitudes. Pairing amplitudes for the ferromagnetic solutions are found to be attractive for both singlet and triplet, but singlet is found to be more attractive. This is due to the interplay and competition between quantum fluctuations and direct interaction. This result raises the intriguing possibility of switching between singlet and triplet via some symmetry-breaking effect. | 14 | 4 | 1404.5655 |
1404 | 1404.2596_arXiv.txt | The $21\,\textrm{cm}$ line provides a powerful probe of astrophysics and cosmology at high redshifts, but unlocking the potential of this probe requires the robust mitigation of foreground contaminants that are typically several orders of magnitude brighter than the cosmological signal. Recent simulations and observations have shown that the smooth spectral structure of foregrounds combines with instrument chromaticity to contaminate a ``wedge"-shaped region in cylindrical Fourier space. While previous efforts have explored the suppression of foregrounds within this wedge, as well as the avoidance of this highly contaminated region, all such efforts have neglected a rigorous examination of the error statistics associated with the wedge. Using a quadratic estimator formalism applied to the interferometric measurement equation, we provide a framework for such a rigorous analysis (incorporating a fully covariant treatment of errors). Additionally, we find that there are strong error correlations at high spatial wavenumbers that have so far been neglected in sensitivity derivations. These error correlations substantially degrade the sensitivity of arrays relying on contributions from long baselines, compared to what one would estimate assuming uncorrelated errors. | Modern cosmological observations have produced exquisite constraints on both the initial and final conditions of structure formation in our Universe. Initial conditions have now been probed to high significance with a large number of cosmic microwave background experiments \cite{HinshawEtAl2013,Planck}, while at low redshifts, a combination of galaxy surveys and traditional astronomical measurements provide the final conditions \cite{SDSS}. Still missing from these direct observations, however, are the intermediate epochs that bridge the gap between early and late times. For example, despite tremendous recent progress in high-redshift galaxy observations, details regarding the formation of the first luminous objects and their effects on the intergalactic medium (IGM) during the Epoch of Reionization (EoR) remain uncertain. In the next few years, direct observations of the EoR will be made possible by measurements of the redshifted $21\,\textrm{cm}$ hyperfine transition of neutral hydrogen (see, e.g., Refs. \cite{Furlanetto2006,Morales2010,Pritchard2012,AviBook} for reviews). At the relevant redshifts, the intensity field of the $21\,\textrm{cm}$ brightness temperature depends on a rich variety of different astrophysical effects, such as fluctuations in the ionization and spin states of the IGM, as well as cosmological quantities such as the underlying dark matter density field and peculiar velocity gradients. A map of the $21\,\textrm{cm}$ intensity field at redshifts $z \sim 6$ and above would therefore be a rich probe of EoR physics, including the nature of the first luminous sources (such as their typical mass and luminosity scales), their ionizing and heating efficiency, and feedback processes on the IGM, among other effects. Such a mapping can be accomplished in three dimensions, since the spectral nature of a $21\,\textrm{cm}$ measurement provides redshift (and therefore line-of-sight distance) information, while the angular directions are mapped using traditional imaging. Because of this, the $21\,\textrm{cm}$ line allows access to a large fraction of our Universe's comoving volume, potentially allowing futuristic measurements to move beyond astrophysics and into the measurement of fundamental cosmological parameters \cite{McQuinn2006,Bowman2007,Mao2008}. There are currently a number of experiments aimed at mapping the fluctuations of the cosmological $21\,\textrm{cm}$ signal, including the Giant Metrewave Radio Telescope Epoch of Reionization experiment (GMRT-EoR \cite{Paciga2013}), the Low Frequency Array (LOFAR \cite{Yatawatta2013}), the Murchison Widefield Array (MWA \cite{Tingay2013}), and the Donald C. Backer Precision Array for Probing the Epoch of Reionization (PAPER \cite{Parsons2010}). These interferometer arrays have yet to make a positive detection of the cosmological signal, with the primary challenges being foreground contamination and the high sensitivity requirements. To increase sensitivity, these experiments are primarily targeting binned, statistical measures of the brightness temperature field such as the power spectrum. Recent progress has resulted in a number of increasingly stringent upper limits \cite{Paciga2013,Dillon2014,Parsons2013}, and proposed next-generation instruments such as the Hydrogen Epoch of Reionization Array (HERA \cite{Pober2014}) and the Square Kilometer Array (SKA \cite{Mellema2013}) promise to yield extremely high significance measurements. In addition to achieving the required sensitivity, observations targeting the redshifted $21\,\textrm{cm}$ line must also contend with foreground contaminants. In the relevant frequency ranges (roughly $\sim 100$ to $200\,\textrm{MHz}$, corresponding to $z\sim13$ to $6$), there exist a large number of non-cosmological sources of radio emission that contaminate measurements. These include sources such as the diffuse synchrotron radiation from our own Galaxy, as well as extragalactic point sources, whether they are bright and resolved or part of a dim and unresolved background. The brightness temperature of foregrounds is expected to be $10^5$ times greater than theoretical expectations for the amplitude of the cosmological signal. A detection of the reionization power spectrum will therefore be challenging without a robust foreground mitigation strategy. Historically, cosmic microwave background (CMB) experiments have had to deal with similar problems of foreground contamination. However, strategies for foreground cleaning that have been developed for the CMB cannot be directly applied to $21\,\textrm{cm}$ cosmology for two reasons. First, CMB experiments typically operate at higher frequencies, where foregrounds are not as bright. In fact, microwave-frequency foregrounds are subdominant to the CMB away from the Galactic plane. In addition, CMB experiments measure anisotropies over a two-dimensional surface, with different observation frequencies providing consistency-checks and a set of redundant measurements that can be used for foreground isolation. The three-dimensional mapping of the $21\,\textrm{cm}$ line, on the other hand, contains unique cosmological information at every frequency, which makes it more difficult to remove foregrounds in a way that does not result in the loss of cosmological signal \cite{Liu2013}. With CMB techniques unlikely to succeed without modification, a number of alternate foreground mitigation strategies have been suggested for $21\,\textrm{cm}$ cosmology. These include spectral polynomial fitting \cite{Wang2006,Liu2009a,Bowman2009,Liu2009b}, Wiener filtering \cite{Gleser2008}, principal component analyses \cite{Paciga2011,Liu2012,Masui2013,Paciga2013}, non-parametric subtractions \cite{Harker2009,Chapman2012,Chapman2013}, Fourier-mode or delay filtering \cite{Petrovic2011, Parsons2012b}, frequency stacking \cite{Cho2012}, Karhunen-Lo\`{e}ve eigenmode projection \cite{Shaw2014a,Shaw2014b}, and inverse covariance weighting \cite{Liu2011,Dillon2013,Dillon2014}. The vast majority of these approaches rely on the fact that foreground sources are expected to be spectrally smooth, while the cosmological EoR signal is expected to fluctuate rapidly with frequency \cite{Oh2003}. The cosmological signal can therefore be extracted by isolating spectrally smooth components from the data. \begin{figure}[t] \centering \includegraphics[width=.45\textwidth]{EoRwindowCartoon.jpg} \caption{A schematic of the EoR window in the cylindrical $k_\perp k_\parallel$ Fourier plane. At the lowest $k_\perp$, errors increase because of limits on an instrument's field-of-view. High $k_\perp$ modes are probed by the longest baselines of an interferometer array, and the sensitivity drops to zero beyond $k_\perp$ scales corresponding to these baselines. Spectral resolution limits the sensitivity at large $k_\parallel$. The lowest $k_\parallel$ are in principle limited by cosmic variance, but in practice the larger concern is limited bandwidth and the foreground contamination, which intrinsically resides at low $k_\parallel$. As one moves towards higher $k_\perp$, however, the foregrounds leak out to higher $k_\parallel$ in a characteristic shape known as the ``foreground wedge". The remaining parts of the Fourier plane are thermal-noise dominated, allowing (with a large collecting area or a long integration time) a clean measurement of the power spectrum in this ``EoR window".} \label{fig:SchematicWedge} \end{figure} Recently, however, a complication to this simple picture was realized, in what has been colloquially termed the ``foreground wedge". Consider a cylindrically-binned power spectrum measurement, i.e. one where Fourier amplitudes are squared and binned in annuli specified by wavenumbers perpendicular to the line-of-sight $k_\perp$ and wavenumbers parallel to the line-of-sight $k_\parallel$. Because the line-of-sight direction is equivalent to the spectral axis of an interferometer, one might have naively expected smooth spectrum foregrounds to be sequestered to only the lowest $k_\parallel$. However, this neglects the fact that interferometers are inherently chromatic instruments, with a given baseline probing finer spatial scales (higher $k_\perp$) at higher frequencies. This coupling of spectral and spatial information is sometimes coined mode-mixing, and results in the leakage of information from low to high $k_\parallel$. This effect is particularly pronounced at high $k_\perp$, where the modes are typically probed by longer baselines, which are more chromatic. Putting everything together, theoretical studies and simulations \cite{Datta2010,Vedantham2012,Morales2012,Parsons2012b,Trott2012,Thyagarajan2013,Hazelton2013} have shown that foregrounds are expected to leak out of the lowest $k_\parallel$ into a characteristic ``wedge" that is schematically shown in Figure \ref{fig:SchematicWedge}. Observations with PAPER and MWA have confirmed this basic picture \cite{Pober2013}, including its evolution with frequency \cite{Dillon2014}. The foreground wedge is both a blessing and a curse. At the sensitivity levels that have been achieved by current experiments, observations have seen a sharp drop-off in foregrounds beyond the wedge \cite{Pober2013}. Theoretical calculations and simulations have shown that such a drop-off is the natural consequence of geometric limitations \cite{Parsons2012b}, provided the foregrounds are spectrally smooth. If further integration reveals low-level foregrounds that are spectrally unsmooth, their influence will leak beyond what is typically labeled as the edge of the wedge. However, if foregrounds continue to be reasonably smooth, the fact that physical considerations limit the extent of wedge implies that there must exist an ``EoR window": a region in Fourier space that is \emph{a priori} expected to be foreground-free. The existence of the EoR window thus enables a relatively robust foreground avoidance strategy, where a detection of the power spectrum can be made simply by avoiding measurements within the wedge. On the other hand, such a conservative approach forces one to work at higher $k$ than if the chromatic effects had not caused the wedge in the first place. This is unfortunate because the ratio of the cosmological signal to instrumental noise typically peaks at low $k$, which means that if it were possible to work within the wedge (or to at least push back its boundaries a little), one would be able to make higher significance detections of the power spectrum. Indeed, in Ref. \cite{Pober2014} it was suggested that working within the wedge can increase the detection significance anywhere from a factor of two to six (depending on the interferometer's configuration), with corresponding decreases in the error bars on astrophysical parameters. Given the potentially high payoff associated with pushing back the influence of the wedge (or equivalently, enlarging the EoR window), it is important to have a statistically rigorous framework for describing the wedge. In this paper, we provide just such a mathematical framework. Since there already exists an extensive literature on the foreground wedge and the EoR window, it is worth summarizing the ways in which this paper builds upon and extends previous results. Works such as Refs. \cite{Datta2010,Vedantham2012,Morales2012} describe instrumental simulations that take one particular realization of foregrounds and propagate them through a power spectrum estimation pipeline. They therefore only probe the \emph{mean} power spectrum, and not the scatter (i.e. the errors) about this mean. In Ref. \cite{Thyagarajan2013}, a statistical treatment of point source populations was considered. While error bars were computed, off-diagonal error correlations (i.e. covariances) in the final measurements were neglected. Potential error correlations are important particularly because Ref. \cite{Thyagarajan2013} considered the application of various tapering functions to their Fourier transforms, and certain choices can result in significant correlations between Fourier bins. We consider full foreground covariances, a full treatment of instrumental effects (such as having a non-tophat beam), a full treatment of data analysis choices such as tapering functions, and a fully covariant propagation of errors. We build upon Ref. \cite{Trott2012}, which made use of Monte Carlo methods to propagate errors. Our treatment is more analytic, allowing us to capture the large dynamic range needed to accurately compute the error statistics in a measurement where the foregrounds are many orders of magnitude brighter than the signal. This is made computationally feasible by our use of the delay spectrum approach introduced in Ref. \cite{Parsons2012b}, where input frequency spectra are sorted into a set of time-delay $\tau$ modes via a per-baseline Fourier transform. However, unlike in some works where the delays are used as an \emph{approximation} for line-of-sight Fourier modes (an assumption that is only valid for short baselines, as we will discuss in Section \ref{sec:InstrumentResponse}), we use delays strictly as a convenient choice of \emph{basis}. This basis makes it computationally possible for us to deal directly with visibilities in our formalism (bypassing any mapmaking steps), which avoids gridding artifacts in our numerical results. We also take into full account the correlations between partially overlapping baselines, and therefore rigorously treat the possible complications that were highlighted in Ref. \cite{Hazelton2013}. A related treatment pertaining to lower-redshift $21\,\textrm{cm}$ intensity mapping experiments (though focusing less on the details of the wedge) can be found in Ref. \cite{Shaw2014b}. While our fiducial calculations are centered around instruments targeting the EoR, the techniques developed in this paper are equally applicable to cosmological $21\,\textrm{cm}$ at lower (or higher) redshifts. We accomplish our goals by making use of the quadratic estimator formalism, which was adapted for $21\,\textrm{cm}$ power spectrum measurements in Refs. \cite{Liu2011,Dillon2013}, and applied to real data in Ref. \cite{Dillon2014}. However, appropriate ``wedge effects" were not incorporated into the formalism, an omission that we rectify in this paper. Placing everything in the quadratic estimator formalism enables a systematic computation of the aforementioned error statistics, as well as a systematic study of the optimality (or lack thereof) of various power spectrum estimators. In a sequel paper (Ref. \cite{Liu2014b}, henceforth ``Paper II"), we will take advantage of this to examine the extent to which statistical methods can enlarge the EoR window. With our fully covariant treatment, we find that the wedge is not simply a region of large foreground errors and biases, but also as a marker for error correlations: at $k_\perp$ values where the wedge is a dominant effect, the errors tend to be strongly correlated. With strongly correlated errors, the number of independently measurable Fourier modes is reduced, suggesting that previous sensitivity estimates (such as those in Refs. \cite{Parsons2012a,Beardsley2013,Pober2014}) may be overly optimistic, particularly for arrays that make use of long baselines (such as LOFAR or GMRT). In fact, the rewards for working within the wedge may be overrated as a result of this, but of course this cannot be quantified without a rigorous way to compute the error statistics of the wedge---hence the present paper. The rest of this work is organized as follows. In Section \ref{sec:InstrumentResponse} we examine the measurement equation of an interferometer in detail, paying special attention to chromatic effects. This provides a first non-covariant preview of the foreground wedge, which we generalize to an approximate, but fully covariant description in Section \ref{sec:basicEstAnalytic}, following a review of the quadratic estimator formalism in Section \ref{sec:QuadEst}. In Section \ref{sec:basicEstNumerical} we discard the approximations made in Section \ref{sec:basicEstAnalytic} in a full numerical implementation of our formalism. We summarize our conclusions in Section \ref{sec:Conclusions}. Because a large number of mathematical quantities are defined in this paper, we provide dictionaries in Tables \ref{tab:Definitions} and \ref{tab:vectMatrixDefinitions} for the reader's convenience. \begin{table*} \caption{\label{tab:Definitions}Dictionary of scalars and functions. The ``context" column gives equation references, typically either their defining equation or their first appearance in the text.} \begin{ruledtabular} \begin{tabular}{lll} Quantity & Meaning/Definition & Context \\ \hline \multicolumn{3}{l}{Basic quantities}\\ $\mathbf{u}$ & Fourier dual to angular direction $\boldsymbol \theta$ & Eq. \eqref{eq:FourierDef} \\ $\eta$ & Fourier dual to $\nu$ (i.e. spectral wavenumber) & Eq. \eqref{eq:FourierDef} \\ $\tau$ & Delay, i.e. Fourier dual to $\nu$ (or spectral wavenumber) for a single baseline & Eq. \eqref{eq:DelayDef} \\ $V(\mathbf{b}, \nu)$ & Visibility measured by baseline $\mathbf{b}$ at frequency $\nu$ & Eq. \eqref{eq:measEqn} \\ $\widetilde{V} (\mathbf{b}, \tau)$ & Delay-space visibility by baseline $\mathbf{b}$ at delay $\tau$ & Eq. \eqref{eq:DelayDef} \\ \hline \multicolumn{3}{l}{Instrumental parameters}\\ $\mathbf{b}$ & Baseline vector & Eq. \eqref{eq:basicVis} \\ $\boldsymbol \theta_0 $ & Characteristic width of primary beam & Eq. \eqref{eq:basicVis} \\ $A$ & Primary beam function & Eq. \eqref{eq:basicVis} \\ $\widetilde{A}$ & Spatial Fourier transform of primary beam function & Eq. \eqref{eq:FirstAtilde} \\ $\widetilde{A}_{b\parallel} $ & Profile of $\widetilde{A}$ parallel to baseline vector direction & Eq. \eqref{eq:SepPrimaryBeam} \\ $\widetilde{A}_{b\perp} $ & Profile of $\widetilde{A}$ perpendicular to baseline vector direction & Eq. \eqref{eq:SepPrimaryBeam} \\ $B_\textrm{chan} $ & Frequency channel width & Eq. \eqref{eq:measEqn} \\ $\gamma $ & Frequency channel profile & Eq. \eqref{eq:measEqn} \\ $\widetilde{\gamma} $ & Fourier transform of frequency channel profile $\gamma$ & Eq. \eqref{eq:vIntegratedGenVtilde} \\ $B_\textrm{band} $ & Bandwidth corresponding to depth of cosmological volume & Eq. \eqref{eq:DelayDef} \\ $\Omega_\textrm{pp}$ & Integrated beam squared area & Eq. \eqref{eq:Omegapp} \\ $T_\textrm{sys}$ & System temperature & Eq. \eqref{eq:NoiseCovarSingelBl} \\ $t$ & Total integration time & Eq. \eqref{eq:NoiseCovarSingelBl} \\ $n(b)$ & Number baselines & Eq. \eqref{eq:NoiseCovarSingelBl} \\ \hline Sky: && \\ $I(\boldsymbol \theta, \nu)$ & Sky brightness temperature at angle $\boldsymbol \theta$ and frequency $\nu$ & Eq. \eqref{eq:basicVis} \\ $\widetilde{I}(\mathbf{u}, \eta)$ & Fourier transform of the sky temperature at angular wavenumber $\mathbf{u}$ and line-of-sight wavenumber $\eta$ & Eq. \eqref{eq:FourierDef} \\ $P(u,\eta)$ & Cylindrically-binned power spectrum at angular wavenumber $u$ and line-of-sight wavenumber $\eta$ & Eq. \eqref{eq:PowerSpectrumAppendix} \\ $C_\ell^X$ & Angular power spectrum of foregrounds, & Eq. \eqref{eq:PowerSpectrumForegrounds} \\ $\nu_c^X$ & Frequency coherence length of foregrounds & Eq. \eqref{eq:PowerSpectrumForegrounds} \\ & (For both $C_\ell^X$ and $\nu_c^X$, $X=\textrm{diff}$ for diffuse Galactic emission and $X=\textrm{ps}$ for point sources) & \\ \hline Data analysis: & &\\ $\phi$ & Bandpass or tapering function & Eq. \eqref{eq:DelayDef} \\ $\widetilde{\phi}$ & Fourier transform of bandpass/tapering function $\phi$ & Eq. \eqref{eq:shortBlVtilde} \\ $h(\mathbf{u}, \eta; \mathbf{b}, \tau)$ & Visibility response at delay $\tau$ of baseline $\mathbf{b}$ to sky mode on spatial scale $\mathbf{u}$ and spectral scale $\eta$.& Eq. \eqref{eq:hDef} \\ $g(u, \eta; b, \tau)$ & Same as $h(\mathbf{u}, \eta; \mathbf{b}, \tau)$ but integrated over direction on $uv$ plane perpendicular to baseline vector & Eq. \eqref{eq:GeeDef} \end{tabular} \end{ruledtabular} \end{table*} \begin{table*} \caption{\label{tab:vectMatrixDefinitions}Dictionary of vectors and matrices. The quantities shown here are grouped into three categories: those that exist in the vector space of the visibility measurements (indexed, for example, by baseline and delay), those that exist in the vector space of bandpowers (indexed by Fourier wavenumbers), and those that bridge the two spaces. In the column giving the length/dimensions, $N_\textrm{bl}$ denotes the number of baselines, $N_\nu$ the number of frequency channels, and $N_\textrm{bands}$ the number of bins in Fourier space (i.e. the number of bandpowers).} \begin{ruledtabular} \begin{tabular}{lllll} Quantity & Components & Meaning/Definition & Length/Dimensions & Context \\ \hline \multicolumn{5}{l}{Quantities in measurement space}\\ $\mathbf{x}$& $\mathbf{x}_i$ & Serialized data vector of visibilities & $N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:xvectdef} \\ $\mathbf{C}$ & $\mathbf{C}_{ij}$ & Total covariance matrix $\mathbf{C} \equiv \langle \mathbf{x} \mathbf{x}^\dagger \rangle$ & $N_\textrm{bl}N_{\nu} \times N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:C=N+S} \\ $\mathbf{N}$ & $\mathbf{N}_{ij}$ & Noise covariance matrix & $N_\textrm{bl}N_{\nu} \times N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:C=N+S} \\ $\mathbf{S}$ & $\mathbf{S}_{ij}$ & Signal covariance matrix & $N_\textrm{bl}N_{\nu} \times N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:genCovar} \\ \hline \multicolumn{5}{l}{Quantities in power spectrum space}\\ $\mathbf{p}$ & $p_\alpha$ & Serialized vector of true bandpowers (i.e. power & $N_\textrm{bands}$ & Eq. \eqref{eq:CovarDecomp} \\ & & spectrum at various grid points in Fourier space) & & \\ $\widehat{\mathbf{p}}$ & $\widehat{p}_\alpha$ & Estimator for bandpowers derived from measurements & $N_\textrm{bands}$ & Eq. \eqref{eq:GenQuadEst} \\ $\boldsymbol \Sigma$ & $\boldsymbol \Sigma_{\alpha \beta}$ & Error covariance of estimated bandpowers, & $N_\textrm{bands} \times N_\textrm{bands}$ & Eq. \eqref{eq:matrixErrorCovar} \\ & & i.e., $\boldsymbol \Sigma_{\alpha \beta} \equiv \langle \widehat{p}_\alpha \widehat{p}_\beta \rangle - \langle \widehat{p}_\alpha \rangle \langle \widehat{p}_\beta \rangle $ & & \\ $\overline{\boldsymbol \Sigma}$ & $\overline{\boldsymbol \Sigma}_{\alpha \beta}$ & Error correlation of estimated bandpowers & $N_\textrm{bands} \times N_\textrm{bands}$ & Eq. \eqref{eq:ErrorCorr} \\ $\mathbf{W}$ & $\mathbf{W}_{\alpha \beta}$ & Window function matrix & $N_\textrm{bands} \times N_\textrm{bands}$ & Eq. \eqref{eq:EstWindAndBias} \\ $\mathbf{b}$ & $\mathbf{b}_\alpha$ & Power spectrum estimator bias & $N_\textrm{bands}$ & Eq. \eqref{eq:bias} \\ $\mathbf{M}$ & $M_\alpha$ & Power spectrum estimator normalization & $N_\textrm{bands}$ & Eq. \eqref{eq:basicEstEalpha} \\ \hline \multicolumn{5}{l}{Hybrid quantities}\\ $\mathbf{E}_\alpha$ & $(\mathbf{E}_\alpha)_{ij}$ & Estimator matrix for quadratic estimator of & $N_\textrm{bl}N_{\nu} \times N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:GenQuadEst} \\ & & of bandpower $p_\alpha$, i.e., $\widehat{p}_\alpha = \mathbf{x} \mathbf{E}_\alpha \mathbf{x}$ & for each $\alpha = 1$ to $N_\textrm{bands}$ & \\ $\mathbf{C}_{,\alpha}$ & $(\mathbf{C}_{,\alpha})_{ij}$ & Response of total covariance to the $\alpha$th bandpower, & $N_\textrm{bl}N_{\nu} \times N_\textrm{bl}N_{\nu}$ & Eq. \eqref{eq:CcommaAlpha} \\ & & i.e., $\mathbf{C}_{,\alpha} \equiv \partial \mathbf{C} / \partial p_\alpha$ & for each $\alpha = 1$ to $N_\textrm{bands}$ & \\ \end{tabular} \end{ruledtabular} \end{table*} | \label{sec:Conclusions} In any measurement of the redshifted $21\,\textrm{cm}$ power spectrum, foreground contamination is a serious concern. Fortunately, observations and various theoretical studies have shown that despite complications arising from the inherently chromatic nature of an interferometric measurement, smooth spectrum foregrounds occupy a characteristic wedge region in cylindrical $k_\perp k_\parallel$ Fourier space. The complement of this region is expected to be relatively foreground-free, forming an EoR window where measurements might be made. While there exists an extensive literature on the topic, previous studies have typically focused on how the foreground wedge manifests itself in the mean power spectrum signal. However, the same physical effects that cause the wedge in the power spectrum also affect the associated error statistics, such as the error covariance and the window functions. An examination of some of these statistics was performed in Ref. \cite{Trott2012} using Monte Carlo methods. In this paper, we have provided a complementary treatment by deriving a rigorous, fully-covariant mathematical description of the foreground wedge and the EoR window. While our methods require the numerical \emph{evaluation} of some matrix expressions, they differ from previous work in that they do not require numerical \emph{simulations} of interferometric measurements, since the underlying framework is largely analytic. This makes it possible to compute error statistics with very high dynamic range, which is crucial since the foregrounds are expected to dwarf both the instrumental noise and cosmological signal. Our formalism takes advantage of the delay spectrum techniques introduced in Ref. \cite{Parsons2012b} to achieve computational savings, and in fact it is the use of the delay basis that makes our covariant, high dynamic range calculations numerically feasible. However, we re-emphasize that this is merely a choice of basis, and that our results are independent of this choice. This was shown explicitly in Section \ref{sec:basicEstAnalytic}, when we developed a description of the foreground wedge in terms of window functions. Our description decouples the causes of the wedge---which depend only on the chromatic nature of the instrument and the specific form of our power spectrum estimator---from the detailed nature of the foreground emission. Independent of foreground properties, window functions that are centered at high $k_\perp$ will typically develop long tails towards low $k_\parallel$. The wedge then results from the additional assumption that foregrounds are spectrally smooth, so that strong signals from low $k_\parallel$ are transferred to higher $k_\parallel$ by the long tails. Once the window functions have been computed, however, our formalism allows such assumptions to be relaxed. With a fully covariant framework, we are able to track all error correlations in our numerical computations. We find that measurements made at high $k_\perp$ have highly correlated errors, effectively reducing the number of independent measurements that can be made in that part of Fourier space. This is particularly important for sensitivity forecasts that rely heavily on measurements made within the wedge, since the wedge's extent in Fourier space is roughly on the same scale as that of the error correlations. Previous studies have typically neglected error correlations, assuming that errors are independent as long as the spatial Fourier cells are of the same size as an antenna's $uv$ footprint, and the spectral Fourier cells are on the order of $1/ B_\textrm{band}$. Our work suggests that this is likely to be too optimistic an assumption. At the highest $k_\perp$ considered in our numerical computations ($k_\perp = 0.13\,h$Mpc$^{-1}$), for example, error correlations reduce the number of independent modes by approximately a factor of 2. This effect will be even more pronounced at even higher $k_\perp$, which are probed by experiments with extremely long baselines. Since the chromatic effects that caused the wedge are closely related to those that cause error correlations, it will be crucial in future research to address the question of exactly how far the wedge can be pushed back (or equivalently, how much one can expand the EoR window). In Paper II, we use the formalism of this paper to explore statistical methods for enlarging the EoR window \cite{Liu2014b}. In this paper, our goal was to provide a rigorous treatment of the wedge. Previous treatments have typically made different simplifying assumptions. These include neglecting partially redundant baselines, approximating delay modes as $\eta$ modes, making assumptions about baseline length, assuming top-hat primary beams, neglecting binning artifacts, or assuming that errors are uncorrelated on the Fourier plane. Our framework discards all of these approximations simultaneously, and it is gratifying to see that the basic picture of the EoR window as a naturally foreground-free region of Fourier space remains unchanged. This bodes well for foreground avoidance efforts that aim to detect the EoR by working outside the wedge, making it possible for $21\,\textrm{cm}$ cosmology to open a new window into the high redshift universe using only existing data analysis techniques, with even more transformative results possible with further advances that expand the EoR window. | 14 | 4 | 1404.2596 |
1404 | 1404.2075_arXiv.txt | Influence of the meteoroid bombardment on properties of the lunar exosphere has been confirmed. Quick increase in the zenith column density of Na atoms above the lunar north pole on August 13, 2009 at 0--1 UT is detected and explained by numerous collisions of relatively small Perseid meteoroids ($<$1~kg) with the Moon during maximum of the Perseid meteor shower. New stringent upper limits of the column densities for Ca, Ba, and Ti atoms in the lunar exosphere are obtained as $5{\times}10^7$, $2.2{\times}10^6$, and $6.9{\times}10^8$~cm$^{-2}$, respectively. It is found that the content of impact-produced Ca and Al atoms in the lunar exosphere is depleted as compared to that of Na atoms. | \label{Introduction} Spectral lines of atoms of sodium and potassium were discovered in the lunar exosphere by Potter and Morgan (1988). The emission lines of neutral sodium were detected at distances of about 5 lunar radii from the center of the Moon on the sunward side and much fainter emission was detected at distances up to about 15--20 lunar radii on the antisunward side. The typical velocity of Na atoms in the extended lunar coma is about 2 km/s suggesting a high-energy source, such as impact vaporization of the regolith (Mendillo et al., 1991). The distribution of sodium in the lunar exosphere depends on solar zenith angle, suggesting that most sodium atoms are liberated from the lunar surface by solar photons or by solar wind impact, in contrast to a source driven by uniform micrometeoroid bombardment (Flynn and Mendillo, 1993; Mendillo et al., 1993). Properties of the extended lunar sodium exosphere have been explained by a 15\% contribution of sporadic micrometeoroid impact vaporization occurring uniformly over the lunar surface and an 85\% contribution of photon-induced desorption dependent on solar zenith angle over the sunlit hemisphere (Mendillo et al., 1999). In particular, impact vaporization caused by collisions of sporadic meteoroids with the Moon may account for up to 50\% of exospheric Na atoms over the terminator and poles (Sarantos et al., 2010). Increase in the content of sodium atoms in the lunar exosphere is expected during activity of main meteor showers due to increasing intensity of meteoroid bombardment. For example, a small increase in temperature and column density of Na atoms in the lunar exosphere was detected during Leonid 1995 and 1997 showers (Hunten et al., 1998; Verani et al., 1998), but no similar effects were detected during Geminid 1999 and Quadrantid 1999 meteor showers (Barbieri et al., 2001; Verani et al., 2001). A bright Na spot in the lunar orbit was detected after maximum of the Leonid 1998 meteor shower (Smith et al., 1999). Thus, meteoroid impacts may lead to the production of Na atoms which are able to escape the lunar exosphere. \begin{table*} \caption{Parameters of spectral observations of Na atoms in the lunar exosphere on August 12--14, 2009. The values of $N_{zen0}$ are calculated at the assumed temperature of 3000~K and averaged from analysis of both Na (5890 and 5896~\AA) lines. The accuracies of the values of observed intensities $I_{obs}$ and the surface zenith column densities $N_{zen0}$ are given at the 3 sigma level.} \small \begin{tabular}{lllllll} \hline Time of observations & Illuminated & Distance from the & Position angle & Intensity of & Intensity of & $N_{zen0}$(Na) \\ UT & fraction, (\%) & surface, (km) & (deg) & Na D2 line, (R) & Na D1 line, (R) & (cm$^2$)\\ \hline Aug.12, 23:13--23:43 &58.8&90 &18.9&146$\pm$6&68$\pm$6&(8.2$\pm$0.5)${\times}10^8$\\ Aug.12, 23:54--Aug.13, 0:24&58.5&270&18.8&159$\pm$3&77$\pm$3&(1.23$\pm$0.04)${\times}10^9$\\ Aug.13, 0:43--1:13 &58.2&455&18.7&137$\pm$3&64$\pm$3&(1.33$\pm$0.04)${\times}10^9$\\ Aug.13, 23:22--23:52 &48 &90 &15.2&152$\pm$3&74$\pm$3&(8.5$\pm$0.3)${\times}10^8$\\ Aug.13, 23:53--Aug.14, 0:23&47.7&270&15.1&136$\pm$3&66$\pm$3&(1.00$\pm$0.04)${\times}10^9$\\ Aug.14, 0:26--0:56 &47.4&455&15.0& 89$\pm$3&45$\pm$3&(8.7$\pm$0.5)${\times}10^8$\\ \hline \end{tabular} \end{table*} Several other authors reported sudden changes in the properties of the sodium lunar exosphere which may be associated with impacts of meteoroids. Hunten et al. (1991) detected an increase in the column density of Na atoms at 801 south latitude in the lunar exosphere of about 60\% on October 14, 1990 as compared to the observations of October 12 and 13, 1990 while measurements at the equator showed no substantial change. These results were explained by the action of an unknown low-speed meteor shower. Similar quick changes in the column density of Na atoms above the north pole of the Moon on September 18--19, 1995, were reported and explained by impacts of numerous small low-speed meteoroids by Sprague et al. (1998). Sudden significant changes in Na temperature were detected during observations of the lunar poles on April 19 and May 10, 1998, which were interpreted as possible impacts of meteoroids (Sprague et al., 2012). These possible impacts may be associated with Lyrid (maximum on April 22) and $\eta$-Aquarid (maximum on May 6) meteor showers. Thus, the column density of Na atoms in the lunar exosphere at the poles varies significantly, but the nature of this variability remains poorly understood. The atoms of refractory elements, such as Ca, Mg, Al, and Fe, have not been detected in the lunar exosphere yet, while Mg and Ca atoms in the Hermean exosphere have, and their presence is explained by the action of high-energy mechanisms such as solar wind sputtering (Sarantos et al., 2011; Burger et al., 2012). Several attempts to detect atoms of refractory elements in the lunar exosphere have been performed (Flynn and Stern, 1996; Stern et al., 1997; Halekas et al., 2013; Cook et al., 2013), but all these observations were performed in the absence of main meteor showers. For this reason, meteoroid impacts were not considered in these papers as a source of the lunar exosphere. The search for lines of refractory elements above the north pole of the Moon during the Perseid 2009 meteor shower was performed by Churyumov et al. (2012), and preliminary results of these observations were reported. The mass of impacted Perseid meteoroids was estimated as 15 kg. The upper limits of Ca, Ba, and Ti atoms in the lunar exosphere were estimated using a simple barometric model as $1.6{\times}10^7$, $7.4{\times}10^5$, and $1.2{\times}10^7$cm$^{-2}$, respectively. However, in this paper we use the more appropriate model of Chamberlain (1963) for the study of exospheric atoms. In addition, we also include a geometric factor in the data analysis, and re-analyze original observational data against the early paper of Churyumov et al. (2012). | Increase in the zenith surface column density of Na atoms in the lunar exosphere on August 13, 2009 at 0--1 UT of about 40% detected while observations of August 13/14, 2009 show the almost constant surface column density of Na atoms estimated from all three recorded spectra. Our observations can be better explained by numerous collisions of small Perseid meteoroids ($<1$~kg) with the Moon (the mass flux is about 27~kg/h) with respect of single impact hypothesis (mass of unique meteoroid of about 20~kg). New stringent upper limits of Ca, Ba, and Ti column densities in the lunar exosphere are obtained. Lower limits of depletion factors of Ca and Al atoms in the lunar exosphere during maximum of the Perseid meteor shower are estimated for the first time and explained by formation of silicate dust particles during cooling of an expanded impact-produced cloud. Our observations confirm influence of the meteoroid bombardment on the properties of the lunar exosphere. More observations of the lunar exosphere during activity of meteor showers are required for deeper understanding of these phenomena. | 14 | 4 | 1404.2075 |
1404 | 1404.5463_arXiv.txt | The formation of high-mass stars is usually accompanied by powerful protostellar outflows. Such high-mass outflows are not simply scaled-up versions of their lower-mass counterparts, since observations suggest that the collimation degree degrades with stellar mass. Theoretically, the origins of massive outflows remain open to question because radiative feedback and fragmentation of the accretion flow around the most massive stars, with $M > 15$~M$_\odot$, may impede the driving of magnetic disk winds. We here present a three-dimensional simulation of the early stages of core fragmentation and massive star formation that includes a subgrid-scale model for protostellar outflows. We find that stars that form in a common accretion flow tend to have aligned outflow axes, so that the individual jets of multiple stars can combine to form a collective outflow. We compare our simulation to observations with synthetic H$_2$ and CO observations and find that the morphology and kinematics of such a collective outflow resembles some observed massive outflows, such as Cepheus~A and DR~21. We finally compare physical quantities derived from simulated observations of our models to the actual values in the models to examine the reliability of standard methods for deriving physical quantities, demonstrating that those methods indeed recover the actual values to within a factor of 2--3. | \label{sec:intro} Molecular outflows in high-mass star-forming regions appear to differ from those around low-mass stars not just in their strength, but also in their lack of collimation. \citet{beuthshep05} suggest that more massive stars appear to have less collimated outflows, although conclusive evidence for this awaits high-resolution observations of massive star-forming regions by ALMA. The nearest massive star-forming regions with strong outflows, such as DR~21 \citep{davis1996} and Cepheus~A \citep{narayanan1996} are observed to have complex jet structures rather than the single, well-collimated jets typically observed from low-mass stars. Our published study of the interplay of magnetic fields and self-gravity in the accretion flow around stars with masses $M_* > 15$~M$_\odot$, which we call high-mass stars in subsequent discussion, suggests that accretion disks around such stars may be vulnerable to disruption by gravitational torques from their own accretion flows, preventing them from driving magnetocentrifugal jets at all \citep{petersetal11a}. However, we determined that observed outflows have substantially higher momentum than either the large-scale magnetic tower jet driven by the accretion flow \citep{petersetal11a}, or the ionization-driven outflow \citep{petersetal12b}. \citet{zinnyork07} review the theoretical controversy over whether massive stars form by monolithic collapse in isolated cores or accretion from a cluster environment containing many other stars. Multiple low- and intermediate-mass stars form in the accretion flow onto high-mass stars in analytic \citep{krattmatz06} and numerical models \citep{bonetal04,smithetal09,petersetal10a,petersetal10c,wangetal10}. The exceptions to this are numerical models starting with initial core masses of 100--300~M$_{\odot}$ rather than 1000~M$_{\odot}$ \citep{krumkleinmckee07,comm11,myersetal13}, particularly with strongly centrally concentrated initial density profiles, which \citet{girietal11} show to suppress fragmentation regardless of other physical processes. Indeed, the lack of isolated young O stars has long been noticed observationally \citep{gies1987,gvaramadze2012}, although whether every OB star formed in a group does remain contested \citep{oey2012}. \citet{petersetal12b} suggested that the combined effect of jets driven from the inner disks of the secondary stars could explain the observed properties of outflows from massive-star forming regions, even absent an outflow from the most luminous and massive star. Jets generally can be approximated to have power proportional to the masses of their stellar sources \citep{matzner2002}. In a region with a \citet{salpeter55} initial mass function, the cumulative jet power will thus be proportional to the cumulative mass, which is indeed dominated by the low-mass stars regardless of whether the highest-mass stars actually have a jet or not. The collective action of the intermediate-mass stars should begin to dominate the outflow from the region even before enough mass has accumulated on the most massive object for it to begin emitting significant amounts of ionizing radiation, or for all of the low-mass stars to have formed: \citet{petersetal10a} show that the lower-mass stars form later than the more massive ones in their simulations. The combined action of jets was first simulated in a massive collapsing region by \citet{linaka06} and \citet{nakali07}. This work was extended by \citet{wangetal10}, who coupled the outflow momentum feedback to sink particles. \citet{cunn11} and \citet{krum12} presented models with combined radiative and outflow feedback. These studies included decaying turbulence in the collapsing gas, but neglected rotation, thus maximizing the support provided by the outflows against collapse by minimizing their alignment. However, simulations of massive star formation including turbulence show the angular momentum vectors of protostars in dense groups are initially closely correlated on length scales of a few tenths of a pc, and only become more randomly oriented as the objects grow in mass and accrete more distant material \citep{jappkless04,fisher04}. Even a model including not only turbulence, but also a better treatment of radiative heating, and higher spatial resolution than our model also finds such almost planar structures, extending at least to the size of the star forming region in our model of 1500~AU \citep[see Fig.\ 2 of][]{krumetal07}. We describe a numerical simulation of this early stage of outflow driving in an initially rotating cloud with turbulence only induced by its own gravitational instability. Since we do not include background sources of turbulence, our core collapses to the center and builds up a rotationally-flattened structure there, in which the entire stellar group forms, thus minimizing the support, but maximizing outflow alignment. We expect that more realistic turbulent initial conditions would lead to fragmentation of our cloud on large scales and the formation of several collapsing regions with globally misaligned angular momentum vectors, but at least initially aligned protostellar spin axes in the densest star-forming regions. There is observational evidence for outflow alignment from well-separated protostars with distinct outflows on scales of a few tenths of a pc in the vicinity of DR21 \citep{davis2007} and in Source~G of W49A \citep{nsmithetal09}, which has suggested that they formed from the same flattened rotating cloud. We compare the result to the observed outflows from the nearby young, massive star-forming regions Cepheus~A and DR~21. In later work, we will describe the subsequent development of the H~{\sc ii} region within the outflow. In Sec.\ 2 we describe our numerical model, and in Sec.\ 3 we describe results, including simulated observations. In Sec.\ 4, we compare our results to the observations in order to determine the consistency of our scenario with available evidence. | 14 | 4 | 1404.5463 |
|
1404 | 1404.2844_arXiv.txt | Spectroscopic$+$photometric redshifts, stellar mass estimates, and rest-frame colors from the 3D-HST survey are combined with structural parameter measurements from CANDELS imaging to determine the galaxy size-mass distribution over the redshift range $0<z<3$. Separating early- and late-type galaxies on the basis of star-formation activity, we confirm that early-type galaxies are on average smaller than late-type galaxies at all redshifts, and we find a significantly different rate of average size evolution at fixed galaxy mass, with fast evolution for the early-type population, $\reff\propto (1+z)^{-1.48}$, and moderate evolution for the late-type population, $\reff\propto (1+z)^{-0.75}$. The large sample size and dynamic range in both galaxy mass and redshift, in combination with the high fidelity of our measurements due to the extensive use of spectroscopic data, not only fortify previous results, but also enable us to probe beyond simple average galaxy size measurements. At all redshifts the slope of the size-mass relation is shallow, $\reff\propto M_*^{0.22}$, for late-type galaxies with stellar mass $>3\times \msolc$, and steep, $\reff\propto M_*^{0.75}$, for early-type galaxies with stellar mass $>2\times \msolb$. The intrinsic scatter is $\lesssim$0.2 dex for all galaxy types and redshifts. For late-type galaxies, the logarithmic size distribution is not symmetric but is skewed toward small sizes: at all redshifts and masses a tail of small late-type galaxies exists that overlaps in size with the early-type galaxy population. The number density of massive ($\sim \msola$), compact ($\reff < 2$kpc) early-type galaxies increases from $z=3$ to $z=1.5-2$ and then strongly decreases at later cosmic times. | \label{sec:intro} The size distribution of the stellar bodies of galaxies, and its evolution with cosmic time, provides important clues about the assembly history of galaxies and the relationship with their dark matter halos. The two main classes of galaxies, early and late types, show very different dependencies between size and stellar mass \citep{shen03}. The weak dependence between size and mass for late-type galaxies implies that the high-mass late types, on average, have higher surface mass densities than low-mass late types. In contrast, early types show a more complex relationship between stellar mass and density, with the density peaking for systems with stellar masses around $M_*\sim 4\times 10^{10}~\msol$ and decreasing toward both lower and higher masses, as reflected in the classical \citet{kormendy77} relation. This fundamental difference does not depend on whether classification of early and late types is based on star formation activity, bulge dominance (S\'ersic index), or visual inspection, and it implies that the two types have very different evolutionary and assembly histories. In this paper we present the evolution of the size-mass distribution up to $z=3$ on the basis of spectroscopy and multiwavelength photometry from the 3D-HST survey \citep{brammer12a} and \emph{HST}/WFC3 imaging from CANDELS \citep{grogin11,koekemoer11}. Angular galaxy sizes are measured from the CANDELS imaging as described by \citet{vanderwel12} and the \emph{HST}/WFC3 grism observations from 3D-HST provide spectroscopic confirmation and redshifts for a large fraction of the sample, considerably strengthening -- with respect to previous studies -- the fidelity of estimates for stellar masses and rest-frame photometric properties. So far, most of the previous studies have focused on the evolution of average galaxy sizes of the high-mass end of the distribution ($\gtrsim 5\times \msolb$). Enabled by both the improved data quality and a fivefold increase in sample size, we can now, for the first time, describe the size distribution of galaxies across redshift. \subsection{Size Evolution of Late-type Galaxies}\label{sec:introlate} Tracing the evolution of the size distribution with redshift allows us to test the most basic elements in our theory of galaxy formation. The zeroth-order expectation is that disk scale lengths evolve fast, approximately as the inverse of the Hubble parameter \citep{mo98}, and early and recent work on the average sizes of Lyman break galaxies (LBGs) at high redshifts ($z\sim 2-6$) roughly agree with this expectation for a $\Lambda$CDM cosmology: \citet{giavalisco96}, \citet{ferguson04}, \citet{oesch10}, and \citet{mosleh12} all find rapid size evolution with redshift: $\reff\propto (1+z)^{\beta=-1.1}$. In contrast, the average size at a given stellar mass of the population of disk-dominated galaxies evolves slowly at late times ($z\lesssim 1$) and has been reported to evolve slowly as measured at fixed galaxy mass ($\beta = -0.2$) or not at all \citep{lilly98, ravindranath04, barden05}. The implication would be that the evolution of the disk galaxy population is decoupled from the evolution of the dark matter halo population. One fundamental difference between the results of LBGs and lower-redshift disk galaxies is the rest-frame wavelength at which the sizes are measured: the rest-frame UV light seen for LBGs originates from young stars that may be, and are generally expected to be, distributed differently than bulk of the baryonic and stellar mass, not to mention the consequences of extinction. The advent of ground-based near-infrared imaging surveys helped to bridge the $z<1$ and $z>2$ regimes by enabling size measurements in a consistent manner at a fixed rest-frame wavelength. Early results suggested slow evolution for late-type galaxies up to $z\sim 3$ \citep{trujillo06a}, but the uncertainties at $z>1$ were such that evolution in that regime was not strongly constrained. Later ground-based work pointed at faster evolution at a fixed galaxy mass: \citet{franx08} found $\beta=-0.6$ and \citet{williams10} found $\beta=-0.9$, but precise constraints at $z>1.5$ remained elusive and the apparent tension between the $z\lesssim 1$ work and the near-infrared at $z\sim 1.5$ unaddressed. Several \emph{HST}/NICMOS-based studies of the morphology and structure of massive $z\sim 2$ galaxies in the rest-frame optical eventually led to mostly converged results, with $\beta\sim -0.8$ for massive ($\sim \msola$), star-forming galaxies from $z\sim 2.5$ to the present \citep{toft07,buitrago08,kriek09b}. So far it has remained unclear as to whether the difference with the significantly faster evolution for LBG galaxies ($\beta\sim -1.1$) is caused by morphological $K$ corrections, the difference in mass (the typical LBG has $M_*\sim \msolb$), or physical changes with redshift. In addition, the difference with the previously mentioned studies at $z<1$ \citep{lilly98, ravindranath04, barden05} remains unexplained. Improving the measurement of $\beta$ and its mass dependence is crucial in order take the next step toward understanding disk galaxy formation. In this paper we will address these issues and describe the full size-mass distribution of high-redshift galaxies over a broad range in galaxy mass and redshift. We will \begin{itemize} \item{measure the evolution of the slope of the size-mass relation;} \item{present the size distribution as a function of stellar mass and redshift;} \item{provide a consistent comparison with UV-selected, high-redshift samples.} \end{itemize} \subsection{Size Evolution of Early-type Galaxies}\label{sec:introearly} Over the past five years, more attention has been bestowed on the size evolution of early-type galaxies than on the size evolution of late-type, star-forming galaxies. Interest in the topic was initiated by reports that $z\sim 1.5$ early-type galaxies have remarkably small sizes in \emph{HST}-based rest-frame UV imaging \citep{daddi05,trujillo07} and ground-based near-infrared imaging \citep{trujillo06b}. NICMOS imaging presented by \citet{zirm07}, \citet{toft07}, \citet{stockton08}, and \citet{mcgrath08} provided space-based, rest-frame optical size measurements that strengthened the evidence for rapid size evolution ($\beta= -1$ or faster) as measured at a fixed galaxy mass ($\sim\msola$). This notion became firmly established through larger samples \citep{buitrago08} and the first spectroscopic samples \citep{vandokkum08, damjanov11}. Concerns regarding gross overestimates of the stellar mass content of the compact early-type galaxies were alleviated by dynamical mass estimates of such galaxies at $z\gtrsim 1$ \citep{vanderwel08,cimatti08,newman10,vandesande11,toft12,vandesande13,bezanson13,belli14}, and the analysis by \citet{szomoru10} of ultra-deep imaging of a single compact galaxy has demonstrated the absence of low-surface brightness wings that could have been missed by shallower imaging. Increases in sample size and dynamic range in stellar mass have constrained the average size evolution of early-type galaxies with stellar masses $>5\times \msolb$ to $\beta\sim -1.3$ up to $z=2.5$, with no evidence for a change in the slope of the relation over this mass range \citep{newman12}. The steepness of the relation combined with the non-negligible scatter accommodates observations that early-type galaxies display a large range in size at $z>1$ \citep[e.g.,][]{mancini10,saracco11}. While the observational results have largely converged, the interpretation is still debated. Some authors have considered the average increase in size over time as being due to the addition of new, larger early-type galaxies. While some argue that this cannot reproduce the observations \citep{vanderwel09a}, others argue that a population of compact early-type galaxies (with sizes $\reff\lesssim 2$kpc) exists within present-day clusters, with a number density comparable to that of higher-redshift early-type galaxies \citep{valentinuzzi10,poggianti13}; tension with the absence of such galaxies in the Sloan Digital Sky Survey (SDSS) remains \citep{trujillo09, taylor10}. The crucial observational test is to trace the evolution of the number density of early-type galaxies as a function of their size. \citet{cassata11}, \citet{szomoru12}, and \citet{newman12} show strong evolution in the number density of small galaxies at $0<z<2.5$, while \citet{carollo13} claim no evolution at $0<z<1$. Our use of 5 fields addresses the issue of field-to-field variations that may affect the aforementioned studies based on smaller samples, and it extends the dynamic redshift range of the Carollo et al.~sample. The leading explanation for the size growth of individual galaxies is accretion and tidal disruption of satellite galaxies that gradually build up the outer parts. For this process, the change in size is large compared with the increase in mass: $\Delta \reff \propto \Delta M_*^2$ \citep[e.g.,][]{bezanson09,hopkins09}. This analytical prediction based on conservation of binding energy has been tested through numerical simulations \citep[e.g.,][]{naab09,oser12, bedorf13}. The analytically predicted and simulated evolution in the increased surface mass density at large radii is, in fact, observed \citep{vandokkum10}; in addition, the central stellar density shows little evolution \citep{bezanson09,hopkins09,vandokkum10}, which is also consistent with a minor merger scenario. In other words, the observations show that there is no need, and little room, to physically expand a galaxy by displacing large numbers of stars to large radii through rapid changes in the central potential, as suggested by \citet{fan10}. A possible challenge to the minor merger scenario is posed by the lack of strong evolution in the slope of the mass density profile seen in lensing galaxies \citep{sonnenfeld13}. Until recently, the size evolution of late- and early-type galaxies was usually discussed separately and treated as different topics. However, in order to understand the joint evolution of these classes, one has to take into account the continuous transition of late-type to early-type galaxies seen in particular in the stellar mass range of $\msolb$ to $\msola$ \citep[e.g.,][]{bell04,faber07,brown07,ilbert10,brammer11,buitrago13,muzzin13}. The star-forming progenitors of the small early-type galaxies are now looked for and plausibly identified \citep[e.g.,][]{whitaker11, barro13, barro14, toft14}, but the evolutionary path of the transitioning galaxies has not been fully mapped out. In this paper, regarding the evolution of early-type galaxies, we will \begin{itemize} \item{present the distribution of sizes as a function of stellar mass and redshift, jointly with those of late-type galaxies; and} \item{show the evolution of the number density of early-type galaxies as a function of size.} \end{itemize} After describing the data and sample selection in Section \ref{sec:data} we present and analyze size distributions as a function of redshift and galaxy mass in Section \ref{sec:distr}. We compare our results with previous studies in Section \ref{sec:comparison} and then discuss the implications of our findings in Section \ref{sec:discussion}. We assume the cosmological parameters $(\Omega_{\rm{M}},~\Omega_{\Lambda},~h) = (0.27,~0.73,~0.71)$. Finally, we use AB magnitudes and the \citet{chabrier03} stellar initial mass function. \begin{figure*}[t] \epsscale{1} \plotone{UVJ_mosaic.pdf} \caption{Rest-frame $U-V$ vs.~$V-J$ color distribution for six redshift bins (each 0.5 wide). The two distinct classes of quiescent and star-forming galaxies are separated by the indicated selection criteria to define our early- and late-type galaxy samples.} \label{uvj} \end{figure*} \begin{figure*}[t] \epsscale{1} \plotone{M_mag_mosaic.pdf} \caption{Observed $\mh$ magnitude vs.~stellar mass in six redshift bins. The color coding represents the rest-frame $U-V$ colors, ranging from $U-V=0$ (blue) to $U-V=2$ (red). The horizontal dashed lines indicate the limit ($\mh=24.5$) down to which we can determine sizes with good fidelity. The vertical lines illustrate the resulting mass completeness limits for blue ($U-V=0.5$) and red ($U-V=2.0$) galaxies, respectively. See Section \ref{sec:sample} for further details.} \label{mmag} \end{figure*} \begin{figure*}[t] \epsscale{1} \plotone{M_UV.pdf} \caption{Rest-frame $U-V$ color vs. stellar mass in six redshift bins. Early-type galaxies, defined as illustrated in Figure \ref{uvj}, are shown in red and late types in blue. A clearly defined red sequence is seen up to $z=3$, with an increased prevalence of dusty late-type galaxies toward higher redshifts.} \label{muv} \end{figure*} \begin{figure}[t] \epsscale{1.2} \plotone{Mgrad.pdf} \caption{Wavelength dependence of $\reff$ in bins of stellar mass and redshift; the latter is indicated by the color coding. Late-type galaxies, as defined in Figure \ref{uvj}, with robust size measurements from ACS/F814W, WFC3/F125W and WFC3/F160W imaging are included (see text for details). Generally, sizes are smaller at longer wavelengths, that is, late-type galaxies are bluer in the outer parts. Moreover, this gradient is stronger for more massive galaxies at all redshifts, and the gradient decreases with redshift, at the same rate for all masses. The dotted lines represent the parameterization given in Equation~(\ref{eq:grad}) that we use to correct our size measurements of late-type galaxies.} \label{mgrad} \end{figure} \begin{figure*}[t] \epsscale{1} \plotone{MR.pdf} \caption{Size-stellar mass distribution of late- and early-type galaxies (same symbols as in Figure \ref{mmag}). A typical $1\sigma$~error bar for individual objects in the higher-redshift bins is shown in the bottom-right panel. The lines indicate model fits to the early- and late-type galaxies as described in Section \ref{sec:fitsize}. The dashed lines, which are identical in each panel, represent the model fits to the galaxies at redshifts $0<z<0.5$. The solid lines represent fits to the higher-redshift samples. The mass ranges used in the fits are indicated by the extent of the lines in the horizontal direction. Strong evolution in the intercept of the size-mass relation is seen for early-type galaxies and moderate evolution is seen for the late-type galaxies (also see Figure \ref{parevol}). There is no significant evidence for evolution in the slope (also see Figure \ref{parevol}). The parameters of the fits shown here are given in Table \ref{tab:fitsize}.} \label{mr} \end{figure*} \begin{figure*}[t] \epsscale{0.38} \plotone{z_r.pdf} \plotone{z_sl.pdf} \plotone{z_sc.pdf} \caption{Parameterized redshift evolution of the size-mass relation, from the power law model fits shown in Figure \ref{mr}. The left-hand panel shows the evolution of the intercept, or the size evolution at fixed stellar mass of $5\times 10^{10}~\msol$. Strong evolution is seen for high-mass early-type galaxies and moderate evolution is seen for low-mass early types and for late-type galaxies. The middle and right-hand panels show the evolution of the slope and intrinsic (model) scatter of the size-mass relation, either with little or no evidence for changes with redshift. The open symbols represent the observed scatter: these measurements do not take measurement uncertainties and contamination into account. The fitting parameters shown in this figure are given in Table \ref{tab:fitsize}.} \label{parevol} \end{figure*} \begin{figure}[t] \epsscale{1.2} \plotone{z_r_res.pdf} \caption{Evolution-corrected average sizes at $M_*=5\times\msolb$ for late-type galaxies (top panel, in blue) and early-type galaxies (bottom panel, in red). The values shown here are the values shown in the left-hand panel of Figure \ref{parevol}, divided by $(1+z)^{\beta_z}$ as indicated on the y-axis. The residuals from the best-fitting $(1+z)^{\beta_z}$ law indicate that parameterizing the evolution as a function of the Hubble parameter ($\reff\propto h(z)^{\beta_H}$) may provide a more accurate description of the late-type galaxies. See Section \ref{sec:medsize} for further discussion.} \label{zr_res} \end{figure} \begin{figure*}[t] \epsscale{1.2} \plotone{MRh.pdf} \caption{Median (points) and 16th and 84th percentiles (lines) of the size-stellar mass distributions shown in Figure \ref{mr}. The scatter in $\reff$ does not strongly depend on galaxy mass. Deviations from the power law form of the size-mass relation are clear for massive late-type galaxies and for low-mass early-type galaxies. Note that here we do not account for contamination (misclassified early- and late-type galaxies).} \label{mrh} \end{figure*} \begin{figure*}[t] \epsscale{.55} \plotone{z_Rm_p.pdf} \plotone{z_Rm_sf.pdf} \caption{Median size as a function of stellar mass and redshift for early-type galaxies (left) and late-type galaxies (right). SDSS data points based on \citep{guo09} are shown as open points. Fits to the median sizes of the form $\reff / \rm{kpc} = B_z(1+z)^{\beta_z}$ and $ B_H(H(z)/H_0)^{\beta_H}$ are shown by dotted and solid lines, respectively. The evolution of the early-type galaxies is independent of mass at $M_*>2\times \msolb$: massive galaxies evolve fast and have a steep size-mass relation at all redshifts, while the relation flattens out at lower masses ($\lesssim 10^{10}~\msol$) and evolves less rapidly. The evolution of the late-type galaxies is overall slower, and does not depend strongly on mass. The low-mass early-type galaxies evolve at the roughly the same pace as the late-type galaxies. The median sizes and fitting results are given in Table \ref{tab:medsize}.} \label{zr} \end{figure*} \begin{figure*}[t] \epsscale{1.2} \plotone{Rhist.pdf} \caption{Size distribution histograms for early- and late-type galaxies as a function of stellar mass (as labeled on the right-hand side) and redshift (as labeled at the top). The number of galaxies is given in units of comoving volume to illustrate the growth of the population over time. The early-type size distributions are fit with Gaussians with a fixed dispersion of 0.16 dex. The late-type size distributions are fit with skewed Gaussians with a fixed dispersion of 0.16 dex and skewness $h_3=-0.15$. The panels with thin lines show samples that are below our mass limit.} \label{rhist} \end{figure*} \begin{figure}[t] \epsscale{1.2} \plotone{Rcumu.pdf} \caption{Cumulative size distributions of $\sim L^*$ early-type galaxies (top) and $\sim L^*$ late-type galaxies (bottom) as a function of redshift. While the number density of both early- and late-type galaxies increases over time, the number density of small galaxies declines, implying that the observed evolution in the mean size is not (solely) driven by the addition of larger galaxies. Individual galaxies must evolve in size.} \label{rc} \end{figure} \begin{figure}[t] \epsscale{1.2} \plotone{lbg.pdf} \caption{Size evolution of galaxies in a narrow (0.3 dex) mass bin around $\msolb$. The black points represent UV-bright galaxies (with $U-V<1$ in the rest-frame), selecting a sample akin to LBGs at high redshift. Their size evolution is fast, consistent with the size evolution of UV-selected samples up to $z=7$ as recently determined by \citet{mosleh12} -- also see \citet{oesch10}. The blue points represent late-type galaxies as defined in this paper (see Figure \ref{uvj}), that is, all star-forming galaxies. The size evolution of those is slower at low redshift, consistent with previous measurements at $z<1$ \citep[here,][]{barden05}.} \label{lbg} \end{figure} \begin{figure}[t] \epsscale{1.2} \plotone{small.pdf} \caption{Number density evolution of compact early-type galaxies. In each redshift bin (with equal co-moving volume) we include early-type galaxies with mass $M_* > 5\times \msolb$ and size $\reff (\rm{kpc}) < (M_*/5\times \msolb)^{0.7}$. That is, the slope of the size-mass relation is taken into account: for $M_* = 5\times \msolb$ the size limit is $\reff = 1$~kpc and for $M_* = \msola$ the size limit is $\reff = 1.6$~kpc. The number density first increases with cosmic time, reaching a plateau at $z\sim1.5-2$, after which it strongly decreases toward the present day. The immediate implication is that individual galaxies must grow in size significantly, most likely through merging.} \label{small} \end{figure} | \label{sec:discussion} \subsection{Evolution of Late-type Galaxies}\label{sec:late} Remarkably, the observed pace of size evolution for late-type galaxies is essentially the same as the evolution of the dark matter halo radius at a fixed mass, $R\propto H(z)^{-2/3}$, but only when halo mass and radius are defined with respect to the critical density. In a $\Lambda$CDM universe, if halo mass is parameterized with respect to matter density or virial density (assuming top-hat collapse), then $\Lambda$ causes strong departures from a power law at late cosmic times. The average evolution between $z=2$ and the present is $R\propto H(z)^{-1.06}$ or $\propto H(z)^{-1.24}$, respectively \citep{peebles80}. The interpretation of such a comparison is not straightforward. However, our novel measurement of the slope and scatter of the size mass relation provides new constraints. The intrinsic scatter in galaxy size remains approximately the same at all redshifts ($\sim 0.16-0.19$~dex, see Figure \ref{parevol}) and is comparable to, but perhaps somewhat smaller than, the scatter of 0.25 dex in the halo spin parameter \citep[e.g.,][]{maccio08}. This strongly suggests that at all redshifts the sizes of late-type galaxies are set by their dark matter halos, and it encourages us to examine the relation between galaxy sizes and halo properties further. The power law fits presented in Section \ref{sec:fitsize} imply that there is very little or no evolution in the slope of the size-mass relation; it remains flat, $\alpha \equiv d \log{\reff}/d \log{M_*}=0.22\pm0.03$, at all redshifts $0<z<3$ (Figure \ref{parevol}, middle panel). As argued by \citet{shen03}, the flat slope suggests that the ratio between galaxy mass and halo mass is not a constant: if it were, the size-mass relation would be steeper ($\alpha=1/3$) than observed. The underlying assumption is that galaxy size is proportional to halo size \citep{kravtsov13}, which we here take to be the case for late-type galaxies. Following \citet{shen03}, we use the observed slope ($\alpha\sim 1/5$) to constrain the galaxy mass-halo mass relation and find $m_g\equiv M_{\rm{gal}}/M_{\rm{halo}}\propto M_{\rm{halo}}^{\gamma\sim 2/3}$. The observation that the slope of the size-mass relation does not evolve with redshift provides a very stringent constraint on the models: unless a combination of factors conspire to keep this slope constant, the most straightforward explanation is that the slope of the relation between galaxy and halo mass ($\gamma$) is similar across the redshift range considered here. Indeed, entirely independent estimates of the relationship between galaxy and halo mass, based on clustering measurements and abundance matching techniques, provide strong evidence that $m_g$ depends on halo mass, similarly so at different redshifts \citep[e.g.,][]{conroy09,moster10,behroozi10,wake11,moster13,behroozi13}. In fact, the most recent studies found that $\gamma=2/3$ for halos in the mass range $M_{\rm{halo}}\sim 10^{11-12}~\msol$, in agreement with what we infer on the basis of the slope of the size-mass relation. In addition, \citet{moster13} and \citet{behroozi13} showed that $m_g$ peaks at a similar halo mass ($\sim 10^{12}~\msol$) at all redshifts $z\lesssim 2$, at around a constant value of $m_g\sim 0.05$. The implication is that $m_g$ does not strongly evolve over the (rather narrow) halo mass range $10^{11-12}~\msol$ \citep{behroozi13b}. It is unclear whether the observed pace of galaxy size evolution ($\reff\propto H(z)^{-2/3}$) implies that $\reff$ evolves in proportion to $R_{\rm{halo}}$, as may be expected in the case that galaxy size scales with halo size in the present-day universe \citep{kravtsov13}. It may be a coincidence that the observed pace of evolution is the same as that for halo radii with respect to the critical density, and it appears more natural to expect galaxy sizes to scale with halo mass and radius that are defined in terms of matter or virial density. In this spirit, the tendency for late-type galaxies to display rather slower size evolution than expected has been given ample attention in the literature. \citet{somerville08} argued that because halos are less concentrated at high redshift, baryonic disks are larger in proportion to the virial radii of halos, leading to slower size evolution. In addition, \citet{dutton11} showed that accreting, gaseous disks with a simple but self-consistent prescription for star formation lead to similarly slow evolution of the stellar disk scale radius as a result of recycling of gas and radial variations in star formation. In addition, stellar feedback may have a more direct impact on disk sizes as low-angular momentum material is ejected \citep[e.g.,][]{maller02, brook11}. The sizes we measured are the not strictly disk scale lengths, as we sample the whole galaxy, including the bulge. Therefore, bulge formation in late-type galaxies slows down size evolution as parameterized here. Bulge formation can either occur rapidly, through mergers \citep{toomre72} or clump formation and migration in unstable disks \citep{dekel09,ceverino10}, or gradually, through secular evolution driven by non-axisymmetries in the disk potential \citep{kormendy04}. The prediction of any of these scenarios is that the galaxies with higher global S\'ersic indices will have smaller sizes at a given mass. The observation that evolution is faster at $z>2$ and slower at $z<1$ (see Section \ref{sec:complate} and Figure \ref{lbg}), combined with the appearance of redder, more compact galaxies at late cosmic times, suggests that bulge formation plays an important role in the evolution of half-light radii of late-type galaxies. \subsection{Evolution of Massive Early-type Galaxies}\label{sec:early} The co-moving number density of $L^*$ early types has strongly increased over the redshift range examined in this paper ($0<z<3$), as was shown by, e.g., \citet{bell04, faber07, brown07, ilbert10, brammer11, buitrago13, muzzin13}. Here this is illustrated in Figures \ref{rhist} and \ref{rc}. Naturally, the progenitors of the newly formed early-type galaxies must be looked for among the star-forming, late-type population. The skewed size distribution of late types toward small sizes (see Section \ref{sec:skew} and Figure \ref{rhist}) points at the existence of a population of small late-type galaxies that span the entire size range seen for early-type galaxies. Figure \ref{rc} illustrates that this is the case over essentially the entire redshift range probed by our sample. The tail of small star-forming galaxies shown in Figure \ref{rc} at $z>1.5$ \citep[also see][]{barro13, williams14, barro14} may reflect the intriguing possibility of a scenario in which such small yet massive star-forming galaxies are the immediate progenitors of compact early-type galaxies. Their number density does not rapidly change between $z=3$ and $z=1.5$, whereas the number of early-type galaxies does rapidly increase over that redshift range (see Figure \ref{rc}). This would suggest the continuous emergence of additional small late-type galaxies that represent a transitional phase between the bulk of the late-type population and the early-type population, as recently advocated by \citet{dekel14} on the basis of analytical arguments and simulations. An alternative interpretation is that the star-forming population consists of ``normal'' late types and a population of early-type galaxies that revived their star-formation activity. The simplest implementation of this model, in which these ``frosting'' early types have the same size distribution as the quiescent early types, can be ruled out: the skewed size distribution of late types is not well described by two log-normal distributions centered at the respective peaks of the size distributions for late- and early-type galaxies. In general, the size distribution of the full population of galaxies (early and late types combined) is not observed to be bimodal in the sense that there is no clear gap between two fiducial populations of small and large galaxies, nor can the size distribution be accurately represented by a single Gaussian distribution. More complicated models of the ``frosting'' flavor, in which a large, star-forming disk reassembles to surround a compact, quiescent component, cannot be immediately ruled out. However, such scenarios seem implausible as the implied color and mass-to-light ratio gradients of such galaxies would likely be stronger than observed \citep{wuyts12}. Measurements of the stellar density in the central regions of early- and late-type galaxies can be used to provide further constraints. Whether or not the small late-type galaxies represent a transitional phase, the central idea in the formation of early-type galaxies is that the formation of an early-type galaxy requires the formation of a concentrated stellar body with a high density \citep[e.g.,][]{franx08, bell12}. One possibility is that a substantial amount of material flows to the center under the influence of mergers \citep[e.g.,][]{dimatteo05} or violently unstable disks and clump formation/migration \citep{dekel09,ceverino10,dekel14}. It remains to be seen whether such processes can reproduce the correct stellar density profiles \citep{wuyts10}. As we showed in Section \ref{sec:compearly} and Figures \ref{rc} and \ref{small}, the number density of small, compact early-type galaxies strongly decreases between $z\sim 1.5$ and the present. This immediately implies that early-type galaxies, after they first form as compact, quiescent objects, have to substantially grow in size over time. Combining this with the suggestion that new early-type galaxies likely form out of the smallest late-type galaxies, the implication is that early-type galaxies are the most dense \citep[and disk-like in structure, e.g.,][]{vanderwel11a, bruce12, chang13} immediately after their star formation is truncated. The amount of later evolution in size and density is dictated by the (non-evolving) slope of the size-mass relation and the evolution of its intercept. This naturally fits into the general idea that a gas-rich formation phase is followed by a more quiescent, dissipationless formation phase. The scatter in the size-mass relation of $\sim 0.15-0.20$~dex (see Figure \ref{parevol}) shows that there is some variation in the amount of dissipative and dissipationless formation, yet, the fact that we see little or no evolution in the size scatter, as predicted by \citet{shankar13}, implies that the amount of dissipation integrated over cosmic history does not vary greatly. Some early-type galaxies may have experienced an intensely dissipative phase at early times, while other -- similarly massive -- galaxies may have continued a less intense star-forming phase up until recently. The compact $z>1.5$ early-type galaxies would fall in the former category; the large, massive star-forming galaxies at $z\sim 0.5-1.5$ may be the progenitors of galaxies in the latter. Within this framework, independent evidence for the increase in stellar mass of individual early-type galaxies by a factor of 2 to 3 between $z=2$ and the present \citep{vandokkum10} implies that the growth in size depends on the growth in mass as $\Delta \reff\propto \Delta M_*^{\sim 2}$. This steep dependence is consistent with a merger scenario. Satellite galaxies can be stripped and their stars deposited on large-radius orbits. Direct and stringent constraints on the minor merger rate are difficult to obtain, but it has proved to be difficult to observationally confirm a sufficiently large minor merger rate to explain the observed evolution \citep{newman12}. Mergers among galaxies that occupy the size-mass relation for early-type galaxies, that is, pure dry mergers, may not occur at sufficient rates \citep[e.g.,][]{nipoti12}. Alternatively, mergers between similarly massive galaxies with different sizes can induce large changes in the size-mass distribution of the population. Assuming that the size distribution of progenitors partaking in major mergers is the same as that of the population as a whole, a $\reff=1$~kpc early-type galaxy at $z\sim 2$ will merge, on average, with a late-type galaxy that is 3 times larger. The properties of the merger remnant will depend on the amount of dissipation and the dynamics of the merger, but it is conceivable that the remnant will be much larger than the compact progenitor. A dense inner region will remain in place, and the strong correlation between central density and quiescence implies that the remnant is likely to be quiescent as well. The mass ratio distribution in the merger history of early-type galaxies, and its effect on size evolution, will remain a topic of debate. However, merging can account for, and is arguably required to explain, the disappearance from $z\sim 2$ to the present of disklike structures among $L^*$ early types \citep{vanderwel11a,bruce12,chang13}, and the observation that the most massive galaxies in the present-day universe do not have a disklike structure, but are intrinsically round \citep{vanderwel09b}. A combined analysis of the evolution of size and morphological properties \citep[see, e.g.,][]{huertas13} will aid to simultaneously interpret size growth and disk destruction. The above narrative shows that we have gathered a plausible set of mechanisms that may play a role in explaining the formation and subsequent evolution of early-type galaxies. Despite this, we lack the basis of a simple analytical framework that is similar to our model for disk formation. However, we note that the rapid pace of size evolution is very close to the size evolution expected for halos as defined by their virial density: $R\propto H(z)^{-1.24}$ for halos compares well with $\reff\propto H(z)^{-1.29}$ for massive early-type galaxies. If we assume that these galaxies only grow through the accretion of other halos and their stellar content, then it is perhaps not a coincidence that halos and galaxies both follow the evolutionary path expected for a dissipationless, top-hat collapse scenario. \subsection{Evolution of Low-mass Early-type Galaxies}\label{sec:lowearly} As we noted in Section \ref{sec:medsize} the slope of the size-mass relation for early-type galaxies flattens below stellar mass $\sim \msolb$, and the size evolution is more comparable to that of late types than that of early types (see Figure \ref{zr}). This suggests that there is a population of low-mass early types that may have formed out of late-type galaxies without going through a transitional phase in which high central densities are attained. The stripping of gas from satellite galaxies is a natural explanation for such evolution and can explain the existence of an excess population of early-type galaxies in clusters that have structural properties similar to those of late-type galaxies in the field \citep{vanderwel10}. Satellites are common in this mass range in the present-day universe \citep[e.g.,][]{vandenbosch08}, but not at higher redshifts, lending the stripping scenario more credence on the basis of based the rapid increase in the comoving number density since $z\sim 1-1.5$ (see Figures \ref{mr} and \ref{rhist}). On the other hand, the early types with mass $\lesssim \msolb$ are $\sim 2$ times smaller than equally massive late types. Disk fading may contribute to this difference, but bulge formation and, in general, processes that cause more massive galaxies to transform into early-type galaxies, may play a role in the low-mass regime as well. A model such as that presented by \citet{peng10} can be expanded in order to separately reproduce the size-mass relations for different types of ``quenched'' galaxies. In this paper we present the size-mass distribution of 30,958 galaxies over a large range in mass ($>\msolc$) and redshift ($0<z<3$), distinguishing between early-type and late-type galaxies on the basis of their star-formation activity. Spectroscopic and photometric redshifts, stellar masses, and rest-frame properties are determined by using data from the 3D-HST survey and auxiliary, multiwavelength photometric data sets spanning from the $U$ band to 8~$\mu$m (see Section \ref{sec:data}). Galaxy sizes are measured from CANDELS imaging by single-component S\'ersic profile fits to two-dimensional light distributions, with a correction for (redshift-dependent) color gradients (Section \ref{sec:size}). Consistent with previous results, we find that high-redshift galaxies are substantially smaller than equally massive, present-day counterparts. As is shown in Figures \ref{mr}, \ref{parevol} and \ref{zr}, late-type galaxies are, on average, a factor of $\sim 2$ smaller at $z=2$ than at the present day, whereas for massive early-type galaxies this is a factor of $\sim 4$. We find that the size evolution of late-type galaxies is marginally better described as a function of the redshift-dependent Hubble parameter, $H(z)$, than as a function of the scale factor, $1+z$ (Figure \ref{zr_res}). Average mass-matched sizes of late- and early-type galaxies evolve as $\reff \propto H(z)^{-0.66} \propto (1+z)^{-.75}$ and $\reff\propto H(z)^{-1.29} \propto (1+z)^{-1.48}$, respectively (Figure \ref{parevol} and Table \ref{tab:fitsize}). High-mass late-type galaxies evolve marginally faster than low-mass late-type galaxies (Figure \ref{zr} and Table \ref{tab:medsize}), but the data are consistent with no evolution in the overall slope of the size-mass relation. At all redshifts $z\le 3$ we find that the slope is shallow for late-type galaxies ($\reff\propto M_*^{0.22}$ for galaxies with stellar mass $M_*>3\times 10^9~\msol$) and is steep for early-type galaxies ($\reff\propto M_*^{0.75}$ for galaxies with stellar mass $M_*>2\times 10^{10}~\msol$). The size-mass relation for lower-mass early-type galaxies is more similar to that of late types than that of high-mass early types (Section \ref{sec:lowearly}). Once cross-contamination between the two classes of galaxies and outliers are taken into account (Figure \ref{parevol} and Section \ref{sec:fitsize}), we also find no evidence for evolution in the (intrinsic) size scatter at a fixed galaxy mass. The implications of these results are discussed in Section \ref{sec:discussion}. The data presented here are consistent with essentially most published data sets (Section \ref{sec:comparison}). Because of the sample size and dynamic range in mass and redshift, the immediate implications of the measurements are less ambiguous than was the case for previous studies. In particular, we show in Figure \ref{rc} that the size distribution of $z\sim 2$ early-type galaxies is significantly different from that of any subset of low-redshift galaxies with the same comoving number density; small early-type galaxies, which are typical at $z\sim 2$, do not exist in equal numbers today (Figure \ref{small}) and must therefore undergo strong size evolution in the intervening time. The size-mass distributions from the 3D-HST and CANDELS projects presented here provide a solid framework for galaxy evolution models, and strongly constrain the interplay between structure formation and galaxy formation \citep[e.g.,][]{stringer13}. The steadily evolving intercept of the size-mass relation, in combination with the non-evolving slope and scatter, present tight constraints on how baryons condense and form galaxies at the centers of dark matter halos (e.g., Section \ref{sec:late}). The different assembly mechanisms of early- and late-type galaxies act similarly at all redshifts, as evidenced by the very different, but unchanging slopes of their respective size-mass relations. | 14 | 4 | 1404.2844 |
1404 | 1404.7185_arXiv.txt | We report on deep observations of the extended TeV gamma-ray source MGRO J1908+06 made with the VERITAS very high energy (VHE) gamma-ray observatory. Previously, the TeV emission has been attributed to the pulsar wind nebula (PWN) of the {\it Fermi}-LAT pulsar PSR J1907+0602. We detect MGRO J1908+06 at a significance level of 14 standard deviations (14~$\sigma$) and measure a photon index of $2.20\pm0.10_{\rm stat}\pm0.20_{\rm sys}$. The TeV emission is extended, covering the region near PSR J1907+0602 and also extending towards SNR G40.5--0.5. When fitted with a 2-dimensional Gaussian, the intrinsic extension has a standard deviation of $\sigma_{src} = 0.44^\circ\pm0.02^\circ$. In contrast to other TeV PWNe of similar age in which the TeV spectrum softens with distance from the pulsar, the TeV spectrum measured near the pulsar location is consistent with that measured at a position near the rim of G40.5--0.5, $0.33\arcdeg$ away. | \label{sec:Introduction} MGRO J1908+06 was discovered by the Milagro collaboration during their seven-year survey of the Northern Hemisphere in very high energy (VHE; $E>100$ GeV) gamma rays at a median energy of 20 TeV \citep{Abdo07}. Observations with the High Energy Stereoscopic System (H.E.S.S.) of the location of MGRO J1908+06 were performed as part of a Galactic plane survey \citep{hess_plane}, and after the announcement of the Milagro source, follow-up observations confirmed an extended VHE source, HESS J1908+063 \citep{Aharonian09}. The extended VHE source was also confirmed by observations with VERITAS \citep [Very Energetic Radiation Imaging Telescope Array System;][]{ward}. These results noted the possibility of an association with the nearby shell-type supernova remnant (SNR) G40.5-0.5 \citep{green09}. However, the measured extension of the VHE source, well beyond the boundary of the radio SNR, required either an additional source of gamma rays or the presence of dense molecular matter with which ultra-relativistic particles could interact and produce gamma rays \citep{Aharonian09}. \citet{Abdo09} discovered a radio-quiet gamma-ray pulsar, PSR J1907+0602, in the vicinity of MGRO J1908+06 with the Large Area Telescope (LAT) on the {\it Fermi} satellite. A later investigation by the {\it Fermi}-LAT team concluded that MGRO J1908+06 and HESS J1908+063 were the pulsar wind nebula (PWN) of PSR J1907+0602 \citep{Abdo10b}. (Hereafter, J1908 will be used to generically refer to the VHE source.) Milagro observed emission from J1908 with a flux corresponding to $\sim$80\% of the Crab Nebula flux at 20~TeV and an upper limit on the intrinsic source extension of $2.6^\circ$ \citep{Abdo07}. The lower energy threshold and better angular resolution of H.E.S.S.\ allowed further investigation of the morphology of the source. A fit of a 2D-Gaussian function to the excess map gave a centroid of RA=$286.98^{\circ}$ and Dec=$6.27^{\circ}$ ($l=40^{\circ}23'9''.2\pm2'.4_{\rm stat}$ and $b=-0^{\circ}47'10''.1\pm2'.4_{\rm stat}$ with systematic errors of $20\arcsec$ per axis) and an intrinsic extension of $\sigma_{\rm src}=0.34^{\circ}~^{+0.04}_{-0.03}$ \citep{Aharonian09}. PSR J1907+0602 has a characteristic age of 19.5~kyr and a spin-down power of $3\times10^{36}$ erg/s \citep{Abdo10b}. The time-averaged flux density at 1.51~GHz was 3.4~$\mu$Jy in measurements made with the Arecibo 305~m radio telescope. From the position and the dispersion measure, the distance was estimated to be 3.2~kpc with a nominal error of $20\%$ \citep{Abdo10b}. Chandra observations of the location of PSR J1907+0602 revealed a faint X-ray counterpart with a non-thermal component that is possibly extended \citep{Abdo10b}. \citet{Downes80} estimated the age of SNR G40.5-0.5 to be 20-40 kyr, and its distance is estimated by \citet{yang} to be 3.4 kpc. Both quantities are compatible with the measured properties of PSR J1907+0602; thus the two objects may be physically associated. However, distance estimates using other methods imply a distance of 5.5--8.5 kpc \citep{yang,Downes80} for the SNR. The {\it Fermi}-LAT team used a Very Large Array Galactic Plane Survey \citep[VGPS;][]{Stil} 1420 MHz continuum image to estimate the position of the SNR to be RA=$286.79^\circ$, Dec=$6.50^\circ$, with an angular diameter of $0.43^\circ$. At a distance of 3.2~kpc, the separation between the SNR and the pulsar is then 28~pc \citep{Abdo10b}. If the pulsar is the progenitor of SNR G40.5-0.5, and assuming that the pulsar originated at the center of the SNR, with the characteristic age of 19.5 kyr, the transverse velocity of the pulsar would be 1400 km/s \citep{Abdo10b}. This velocity is $\sim$3 times higher than the typical velocity given to a pulsar due to a supernova ``kick" \citep{slanereview}, but three other pulsars are estimated to have transverse velocities above 1000 km/s: PSR B2224+65 \citep[$\sim1600$~km/s; ][]{pulsar1}, PSR B2011+38 \citep[$\sim1600$~km/s; ][and references therein]{pulsar2}, and PSR B1508+55 \citep[$\sim1100$~km/s; ][]{pulsar3}. Table \ref{tab:sources} summarizes the parameters of PSR J1907+0602 and SNR G40.5--0.5. VERITAS observed J1908 repeatedly during 2007-2012, obtaining a total exposure of 62~hours. Here, we describe this deep TeV observation taken in order to help understand the physical origin of the emission. We describe the observations in Section 2 and analysis in Section 3. We present our results on the morphology and spectrum of the TeV emission in Section 4. We discuss interpretations of the results in Section 5. \begin{deluxetable}{lcc} \tablecaption{Parameters of PSR J1907+0602 and SNR G40.5--0.5 \label{tab:sources}} \tablehead{ Parameter & PSR J1907+0602 & SNR G40.5--0.5} \startdata Right Ascension & $286.978^\circ$ & 286.786$^\circ$ \\ Declination & $6.038^\circ$ & 6.498$^\circ$ \\ Age & 19.5 kyr & 20--40 kyr \\ Distance & 3.2 kpc & 3.4 \\ \enddata \tablecomments{Data are from \citet{Abdo10b} except for G40.5--0.5 distance \citep{yang} and age \citep{Downes80}.} \end{deluxetable} \begin{figure*}[t] \centering {\includegraphics[width=6.0in]{gammasky_axes.eps}} \caption{Map of VHE gamma-ray excess counts in the region near MGRO J1908+06. The color scale shows the excess counts. The black cross shows the VERITAS best fitted position. The three black circles show the regions used for spectral analysis. The white circle shows the size of the VERITAS point spread function (PSF). The red ellipse shows the extent of SNR G40.5--0.5 in radio continuum emission. The black `x' shows the location of PSR J1907+0602. The white cross shows the H.E.S.S. best fitted position and the white contours are from a H.E.S.S.\ excess map at levels of 4, 5, 6, and 7~$\sigma$. The grids on this and all sky maps in this paper are at intervals of $0.5\arcdeg$.} \label{fig:gammamap} \end{figure*} | \label{sec:Discussion} The TeV emission from J1908 is strong in the region near PSR J1907+0602 and also extends towards SNR G40.5--0.5. A key question regarding the nature of the emission is whether it is solely due to a PWN associated with PSR J1907+0602 \citep{Abdo10b} or whether there are additional sources of TeV emission. We estimate a physical size for J1908 of 50~pc, assuming the pulsar distance of 3.2~kpc and adopting twice the $\sigma = 0.44\arcdeg \pm 0.02\arcdeg$ width of the 2D Gaussian fit as the source diameter. Comparison with other TeV PWNe shows that J1908 is larger than other TeV PWNe of similar age (see Figure~6 in \citet{Kargaltsev13}\footnote{We note that the size of 70~pc for HESS J1825-137 quoted by \citet{Kargaltsev13} is a factor of $\sim 2$ larger than the size of 36~pc measured in \citet{Aharonian06b}.}). However, PWN size may depend on environment and interaction with the host SNR, particularly for older PWNe. Also, J1908 could be made consistent with the size-versus-age relation found for other PWNe if the pulsar is assumed to be a factor of two older than the spin-down age. This is within the uncertainties associated with spin-down age estimates \citep{Gaensler00} and would also reduce the proper motion to a more typical value. The 1-10~TeV luminosity of J1908 of $3 \times 10^{34} \rm \, erg \, s^{-1}$ is 1\% of the spin-down power of PSR J1907+0602 of $3 \times 10^{36} \rm \, erg \, s^{-1}$, which gives a TeV gamma-ray efficiency similar to that of other PWNe with pulsars of similar spin-down power \citep{Kargaltsev13}. X-ray observations reveal hard emission from the location of the pulsar that is possibly extended and could be a compact X-ray PWN or possibly a bow shock \citep{Abdo10b,Pandel12}. Most, but not all, TeV PWN with similar pulsar spin-down power are detected in X-rays \citep{Kargaltsev13}. Thus, the gross properties of J1908, except possibly its extent, are compatible with its identification as a TeV PWN powered by PSR J1907+0602. For J1908, we find that the spectrum measured near the SNR boundary at a projected distance of 23~pc from the pulsar is the same, within uncertainties, as the spectrum measured near the pulsar up to energies of at least 10~TeV, see Figure~\ref{fig:spectra}. This may present issues with interpretation of the TeV emission as solely powered by PSR J1907+0602. If PSR J1907+0602 was born at the center of G40.5--0.5, then electrons nearer to the SNR were emitted at earlier times than the electrons near the current pulsar location. These older populations of electrons would have cooled over the lifetime of the pulsar via synchrotron emission and inverse Compton (IC) scattering. Assuming an interstellar magnetic field of 3~$\mu$G and that the seed photons for IC scattering are predominately from the cosmic microwave background (CMB), the cooling time for the 63~TeV electrons needed to produce 10~TeV photons is 12~kyr \citep{Kargaltsev07}. Comparing this cooling time with the pulsar age of $\sim 19.5$~kyr, one would expect significant softening or a fall-off at high energies in the spectrum from the region near the SNR. Thus, the lack of softening is incompatible with the behavior expected due to electron cooling. The discrepancy is worse if the characteristic age underestimates the true age, as needed to reduce the pulsar velocity to a more typical value. Spectral softening is also seen in TeV PWNe containing pulsars without very high proper motions. The TeV PWN most similar to J1908 is HESS J1825--137 with a physical size of $\sim 36$~pc and a parent pulsar with a characteristic age of 21~kyr, but a proper motion of only $0.023\arcsec$/year \citep{Pavlov08}. The TeV spectrum of HESS J1825--137 softens with distance from the pulsar, with a photon index of $\Gamma = 1.83 \pm 0.09$ at the pulsar location and $\Gamma = 2.25 \pm 0.06$ at a projected offset of 24~pc \citep{Aharonian06b}. The change in photon index, $\Delta \Gamma = 0.42 \pm 0.11$ should be compared with that measured above in J1908 at a similar physical separation from the pulsar, $\Delta \Gamma = -0.04 \pm 0.20$. Thus, the lack of spectral softening in J1908 far from the pulsar may again argue against interpreting it as a single PWN, even if the pulsar was not born in G40.5--0.5. A potential counter example is the TeV source TeV J2032+4130 that is a PWN associated with PSR J2032+4127. VERITAS observations of the source show no spectral evolution across its extent (full width at half maximum) of 11~pc \citep{Aliu14}. The TeV emission from TeV J2032+4130 appears to be confined to a void apparent in infrared and radio images. \citet{Aharonian09} note that J1908 lies in a void in molecular line emission integrated over $v_{LSR}$ = 25.3-30.5~km/s corresponding to distances of 1.5--1.8~kpc. The physical size of J1908 would be compatible with that TeV J2032+4130 at this distance. However, \citet{Abdo10b} estimate a distance of $3.2 \pm 0.6$~kpc for PSR J1907+0602 based on the pulsar's dispersion measure and place a lower limit on the distance of 3~kpc based on its X-ray flux. Thus, the distance to PSR J1907+0602 is incompatible with a location in the molecular line emission void unless both pulsar distance estimates are incorrect. Another source, in addition to the PWN associated with PSR J1907+0602, may contribute to the observed VHE emission. Fitting the VHE excess map using a model with two 2D-Gaussians provides only marginal improvement, at the $<95$\% confidence level, over the fit using a single 2D-Gaussian. Thus, we do not claim detection of two separate sources of emission. However, the observed morphology could be produced by two interacting sources or two sources superimposed along the line of sight. Some SNRs produce TeV emission via IC interactions of electrons accelerated in the SNR shock or via interactions of accelerated hadrons with surrounding target material such as molecular clouds. Comparison of the 1420~MHz radio continuum emission from SNR G40.5-0.5 with the TeV emission shows there is little or no radio emission in the regions of strong TeV emission, see Figure~\ref{fig:radiomap}. Figure \ref{fig:comap} shows the VERITAS excess map with contours of CO line emission integrated over $v_{LSR}$ = 45-65 km/s overlaid (the CO map was kindly provided by \citet{yang}). The distance to the CO gas at these velocities, 3.4~kpc, is compatible with the distance to the pulsar and the SNR within the uncertainties. There is CO emission, and thus molecular clouds, at the boundary of the TeV emission, but not coincident with it. In other middle-aged SNRs with TeV emission, the TeV emission is positionally well correlated with CO line and/or radio continuum emission, see Table~\ref{Table:middle_aged_SNR}. Thus, it is unlikely that SNR G40.5-0.5 is the main source of the observed TeV emission. The extended morphology of the TeV emission could be due to interaction of the pulsar wind with molecular clouds or the expanding SNR shell. A shock formed at such an interaction could reaccelerate energetic particles. Alternatively, the molecular clouds may provide seed photons for inverse-Compton scattering on high-energy electrons from the pulsar wind. The dominant seed photons in J1908 are likely microwave photons from the CMB. IR photons from molecular clouds with typical temperatures of 10--50~K are more energetic than the CMB photons, thus, lower-energy electrons with longer cooling times suffice to produce TeV photons. The increased radiation energy density, relative to the CMB, near the observed clouds may brighten the TeV emission near the clouds, causing the extended morphology. In conclusion, J1908 appears to be physically somewhat larger than other TeV PWNe of similar age and the TeV spectrum does not appear to soften with distance from the pulsar, as observed in similar TeV PWNe and expected from electron cooling. There is no strong TeV emission associated with the radio-bright shell of SNR G40.5-0.5. Interaction of the pulsar wind with nearby molecular clouds or the SNR shock could explain the large size and lack of spectral softening in J1908. It is also possible that another PWN, associated with an undetected second pulsar located near the southern edge of SNR G40.5-0.5, could contribute to the VHE emission, but this would require the presence of a second, and as yet undetected, pulsar. | 14 | 4 | 1404.7185 |
1404 | 1404.3502_arXiv.txt | {Spectral differential imaging (SDI) is part of the observing strategy of current and future high-contrast imaging instruments. It aims to reduce the stellar speckles {that prevents} the detection of cool planets {by} using in/out methane-band images. It attenuates the signature of off-axis companions to the star, {such as} angular differential imaging (ADI). However, this attenuation {depends} on the spectral properties of the low-mass companions we are searching for. {The implications of this particularity on {estimating} the detection limits have been poorly explored so far.}} {{We perform an imaging survey to search for cool (T$_{\rm{eff}}$\,$<$\,1\,000--1\,300~K) giant planets at separations as close as 5--10~AU. We also aim to assess {the} sensitivity limits in SDI data taking {the photometric bias into account. This will lead to} a better view of the SDI performance.}} {{We {observed} a selected sample of 16 stars (age $<$ 200~Myr, distance $<$ 25~pc) with the phase-mask coronagraph, SDI, and ADI modes of VLT/NaCo.}} {We do not detect any companions. As for the estimation of the sensitivity limits, we argue that the SDI residual noise cannot be converted into mass limits because it represents a differential flux, {unlike} what is done for single-band images, in which fluxes are measured. This results in degeneracies for the mass limits, which may be {removed} with the use of single-band constraints. We {instead emply} a method {of directly determining} the mass limits and compare the results from a combined processing SDI-ADI (ASDI) and ADI. The SDI flux ratio of a planet is the critical parameter for the ASDI performance at close-in separations ($\lesssim$1$''$). {The survey is sensitive to cool giant planets beyond 10~AU {for 65\% and 30~AU for} 100\% of the sample.}} {{For close-in separations, the optimal regime for SDI corresponds to SDI flux ratios higher than $\sim$2. According to the BT-Settl model, this translates into T$_{\rm{eff}}$\,$\lesssim$\,800~K, which is significantly lower than the methane condensation temperature ($\sim$1300~K).} The methods described here can be applied to the data interpretation of SPHERE. In particular, we expect better performance with the dual-band imager IRDIS, thanks to more suitable filter characteristics and better image quality.} | \label{sec:intro} The search for exoplanets by direct imaging is challenged by very large brightness ratios between stars and planets at short angular separations. Current facilities on large ground-based telescopes or in space allow {adequate contrasts to be reached}, and have revealed a few planetary-mass objects \citep{Marois2008c, Marois2010b, Lagrange2009, Rameau2013b, Kuzuhara2013} either massive ($>$3 Jupiter masses or \mjnospace) and young ($<$100--200~Myr) or with large angular separations ($>$1$''$). Even though some of them are questioned \citep{Kalas2008}, these objects {very} likely represent the top of the giant planet population at long periods. {They are therefore} very important {for understanding the planet's} formation mechanisms. These discoveries have been favored by {longstanding} instrumental developments {such as} adaptive optics (AO) and coronagraphy, but also {by} dedicated observing strategies and post-processing methods like differential imaging. {The} purpose of differential imaging is to attenuate the {stellar speckles which prevent} the detection of faint planets around the star. More precisely, a reference image of the star is built and subtracted {from} the science images. Several kinds of differential imaging have been proposed in the {past} decade. Angular differential imaging {(ADI)} takes advantage of the field rotation occurring in an alt-az telescope \citep{Marois2006a}. Spectral differential imaging {(SDI)} exploits the natural wavelength dependence of a star image \citep{Racine1999}. An extension of this technique consists in using the spectral information in many spectral channels, provided for instance by an integral field spectrometer {\citep{Sparks2002, Thatte2007, Crepp2011, Pueyo2012}}. Polarimetric differential imaging uses differences between the polarimetric fluxes of the star and the planet {or} disk \citep{Kuhn2001}{, but has not permitted any planet detections so far}. Introducing a difference between star and planet properties allows {differentiating} the {unwanted stellar speckles from} the much fainter planet signals. Still, achieving high performance with these methods also {requires good} knowledge of the instrument behavior and biases. Large direct imaging surveys have tentatively constrained the frequency of young giant planets at long periods ($\gtrsim$10--20~AU) to $\sim$10--20\% \citep[{e.g.},][]{Lafreniere2007b, Chauvin2010, Vigan2012, Rameau2013a, Wahhaj2013a, Nielsen2013, Biller2013}. {Typical contrasts} of 10 to 15 magnitudes {have been obtained,} but for separations beyond $\sim$1$''$. Consequently, the occurrence of young giant planets down to a few Jupiter masses was mostly investigated at physical separations from $\sim$10~AU to hundreds of AU, $\beta$~Pictoris~b {being the object detected with the closest separation} \citep[8~AU,][]{Lagrange2010b}. To analyze closer-in, colder, and {less-massive} giant planets, we need to push {the contrast performance further}. For this very purpose, several new-generation imaging instruments {are now ready to start operation{, such as} SPHERE \citep[Spectro-Polarimetric High-contrast Exoplanet REsearch,][]{Beuzit2008} and GPI \citep[Gemini Planet Imager,][]{Macintosh2008}. They} were built to take advantage of several high-contrast imaging techniques{, namely} extreme AO, advanced coronagraphy, {ADI, and SDI.} The evolutionary models extrapolated from stellar mechanisms \citep{Burrows1997, Chabrier2000} predict that {Jovian} planets are very hot when formed; they cool over time and can be relatively bright at young ages\footnote{We note that ``cold-start'' models \citep{Marley2007, Fortney2008, Spiegel2012, Mordasini2012} also predict bright planets at young ages but fainter than those in ``hot-start'' models.}. {SDI \citep{Racine1999}} is intended to take advantage of the presence of a methane absorption band at $\sim$1.6~$\muup$m in the spectra of cool ($\lesssim$1\,300~K) giant planets \citep{Burrows1997, Chabrier2000}, while the star is not expected to contain this chemical element. Thus, this spectral feature provides an efficient tool {for disentangling} stellar speckles from planet signal(s). SDI offers the potential to reach the detection of planets with lower masses than those already discovered by direct imaging. However, spectroscopic observations of a few young giant planets have {only} shown weak absorption by methane in the H band, which could be explained by the low surface gravity of these objects \citep{Barman2011a, Barman2011b, Oppenheimer2013, Konopacky2013}. In practice, SDI also produces a significant attenuation of the planet itself, because the latter is present in the reference image used for the speckle subtraction. {This attenuation has} to be quantified to derive its photometry accurately. To date, only two independent surveys {have been made} using SDI with the Very Large Telescope (VLT){, as well as with} the Multiple Mirror Telescope (MMT) \citep{Biller2007} {and} the Gemini South telescope \citep[][]{Biller2013, Nielsen2013, Wahhaj2013a}, but {they} have not reported {any detections of planetary-mass objects yet}. The non-detection results were exploited to assess the frequency of giant planets at long periods. \citet{Nielsen2008} did not consider the biases introduced by SDI for their analysis, {unlike} \citet{Biller2013}, \citet{Nielsen2013}, and \citet{Wahhaj2013a} for the statistical analysis of the NICI Campaign. {The} observing strategy of the NICI Campaign is based on the complementarity of two observing modes in order to optimize the survey sensitivity: ADI \citep{Marois2006a} {and} the combination of SDI and ADI (ASDI). These two observing modes are not performed simultaneously on the same target, because the spectral filters used are different (large-band and narrow-band{, respectively}). The ADI and ASDI contrast curves presented in \citet{Biller2013}, \citet{Nielsen2013}, and \citet{Wahhaj2013a} are corrected from the attenuation and the artifacts produced by the reduction pipeline except for the SDI part (for the ASDI curves), because the attenuation {depends} on the spectral properties of a planet. Nevertheless, this point is taken into account for the {planet frequency study}. Both ADI and ASDI contrast curves are considered in this analysis, but only the best detection limit is finally used. The results are essentially consistent with the previous surveys. {We} note that \citet{Biller2013} and \citet{Nielsen2013} do not report {any} ASDI mass detection limits for individual targets, because of the particularities of the SDI attenuation. \citet{Wahhaj2013a} present individual mass detection limits combining ADI and ASDI, using the contrast-mass conversion based on evolutionary models. We argue in this work that this method is inadequate for interpreting {the} dual-band imaging data analyzed with SDI-based algorithms. Our arguments are also relevant to IFS data processed with similar techniques \citep{Sparks2002, Crepp2011, Pueyo2012}. In this paper, we present the outcome of a small survey of 16 stars performed with NaCo (Nasmyth Adaptive Optics System and Near-Infrared Imager and Spectrograph), the near-IR AO-assisted camera of the VLT \citep{Rousset2003, Lenzen2003}. Our prime objective is to observe a selected sample of young ($\lesssim$200~Myr) and nearby ($\lesssim$25~pc) stars to search for massive but cool gas giant planets at separations as small as 5--10~AU. For this purpose, we combine state-of-the-art high-contrast imaging techniques similar to those implemented in SPHERE \citep{Beuzit2008}. Our second objective is to address the {problem} of assessing detection limits in SDI data, which is important when it falls to very short angular {separations} ($<$0.5--1$''$), and to determine the condition(s) for which SDI gives the optimal performance. This last topic has not been addressed so far in the literature. {Unlike} the NICI Campaign, we {carried} out the observations of the survey with only one observing mode, ASDI. We consider ASDI and ADI for the reduction and analysis of the same data set, thus {allowing comparison} of the performances of these differential imaging techniques. This paper is designed to focus on the astrophysical exploitation of the survey, based on a simple, straightforward, and robust method {of accounting} for the photometric bias induced by SDI. A subsequent paper will analyze {the details of} the biases of SDI data reduction and {will correctly estimate} the detection performance (Rameau et al., in prep.). The methods and results presented in these papers may serve as a basis for {interpreting} future large surveys to be performed with SPHERE and GPI. We describe the sample selection in Sect.~\ref{sec:targetselection}{, then explain} the observing strategy, the data acquisition, and the reduction pipeline {in Sect.~\ref{sec:obs}}. In Sect.~\ref{sec:sdianalysis}, we explain the {problem} of assessing detection limits in SDI, which requires {a different analysis} from the method usually considered for direct imaging surveys. {In this section, we} also introduce the method we {used for interpreting} our survey. We present {ADI and ASDI detection limits} and carry out a detailed study of the SDI performance in Sect.~\ref{sec:results}. Finally, we discuss the {broad trends} of the survey in Sect.~\ref{sec:discussion}. | We presented the outcome of a high-contrast imaging survey of 16 stars combining the coronagraphic, the spectral differential imaging, and the angular differential {imaging modes} of VLT/NaCo. We did not detect any companion candidates in the reduced images. {We analyzed the sensitivity limits taking {the SDI photometric bias into account} and determined the optimal conditions in terms of SDI performance. The key results are \begin{itemize} \item {By} combining the best detection limits (ADI, ASDI), the survey is sensitive to cool giant planets (T$_{\rm{eff}}$\,$<$\,1\,000--1\,300~K) at projected separations $>$10~AU for 65\% of the sample and $>$30~AU for all targets. We are able to probe the range optimal for ASDI (\teffnospace\,$\lesssim$\,800~K according to the BT-Settl model) for two-thirds of the targets {beyond $\sim$}20~AU. Evolutionary models spanning temperatures $\lesssim$500~K would be needed for the data interpretation of SPHERE and GPI, since we reach this limit for half of the targets. \item {Determination} of the detection limits in ASDI-processed images requires a different analysis than for images processed in single-band differential imaging. In particular, the residual noise level cannot be converted into planet mass through evolutionary models {({regarless of} the considered models), since} it represents a differential flux. This differential flux has to be corrected from the self-subtraction produced by SDI, which depends on the spectral properties assumed for the detectable companions. Thus, detection limits in SDI data should only be considered in terms of physical properties of companions. To derive the detection limits of our survey, we used the signal-to-noise ratio of synthetic planets introduced in the raw data and processed{, as well as the flux predictions of the BT-Settl model}. \item {The} SDI differential flux may be reproduced by several flux couples, hence planet masses. Consequently, the data should also be processed with single-band differential imaging methods to {remove} the degeneracies. \item ASDI can either improve or degrade the sensitivity {regardless of} the angular separation and the star age. For the favorable cases, the gains in detectable planet mass can be as {much as 10\% to 35\%}. The parameter that determines to first order the performance is the SDI flux ratio of the companion. This parameter reaches optimal values when it is $\gtrsim$2 for separations $\lesssim$1$''$. Using BT-Settl, this translates into \teffnospace\,$\sim$\,800~K, which is significantly lower than the methane condensation temperature (1\,300~K). The PSF quality is a second-order factor that modulates the ASDI gain. \end{itemize}} We finally discussed some implications of this work for the data analysis of SPHERE. We expect better performance in particular for ASDI with the dual-band imager IRDIS with respect to NaCo, thanks to more favorable filter characteristics and an extreme-AO system. We envision {a future project of applying} the methods that we developed for this paper {for analyzing} laboratory and commissioning data of IRDIS and IFS. | 14 | 4 | 1404.3502 |
1404 | 1404.4358_arXiv.txt | This study presents the rotational distribution of red giant stars (RGs) in eleven old to intermediate age open clusters. The masses of these stars are all above the Kraft break, so that they lose negligible amounts of their birth angular momentum (AM) during the main sequence evolution. However, they do span a mass range with quite different AM distributions imparted during formation, with the stars less massive than $\sim 1.6$\msun\ arriving on the main sequence with lower rotation rates than the more massive stars. The majority of RGs in this study are slow rotators across the entire red giant branch regardless of mass, supporting the picture that intermediate mass stars rapidly spin down when they evolve off the main sequence and develop convection zones capable of driving a magnetic dynamo. Nevertheless, a small fraction of RGs in open clusters show some level of enhanced rotation, and faster rotators are as common in these clusters as in the field red giant population. Most of these enhanced rotators appear to be red clump stars, which is also true of the underlying stellar sample, while others are clearly RGs that are above or below the clump. In addition to rotational velocities, the radial velocities and membership probabilities of individual stars are also presented. Cluster heliocentric radial velocities for NGC 6005 and Pismis 18 are reported for the first time. | \label{sec:intro} Studies of rotation and angular momentum (AM) evolution are becoming increasingly more sophisticated with the growth of large observational datasets of stellar surface rotation probed by both spectroscopic rotational velocities and photometric rotational periods. Additionally, the exquisite photometric data from \emph{CoRot} \citep{Auvergne:2009en} and \emph{Kepler} \citep{2010Sci...327..977B} has enabled the first large scale studies of the interior rotation of stars other than the Sun. Such internal studies of red giants (RGs) have clearly revealed differential rotation that changes over time. \cite{Deheuvels:2012et} demonstrated that the cores of RGs rotate more rapidly than the surface, and \cite{Mosser:2012dj} found that this core rotation slows significantly sometime during the late stages of the first ascent red giant branch (RGB) evolution. These new observations challenge our understanding of AM transport. \cite{2012A&A...544L...4E} and \cite{2012AN....333..971C} have shown that the physical mechanisms of transporting AM currently invoked in models (e.g., meridional circulation, shear mixing) are insufficient to explain the observed profiles. Understanding the evolution of AM within a star throughout its life may help shine light on the unusually fast surface rotation that is sometimes found for apparently isolated RGs. In the context of RGs, ``fast'' rotation can refer to surface rotations as low as 4 or 5~\kms, because the combination of AM shedding and a growing moment of inertia slows most RGs rotation to projected rotational velocities (\vsini) $\lesssim2$~\kms\ \citep{1996A&A...314..499D}. The fast rotators are relatively rare, found only among a few percent of the field red giant population (e.g., \citealt{1993ApJ...403..708F}, \citealt{2008AJ....135..209M}, and \citealt{2011ApJ...732...39C}). The origin of this unusually high surface rotation is still a matter of debate, in part because some relevant fundamental properties of field RGs are difficult to measure, especially masses. This is because different stages of evolution for different masses can overlap in the Hertzsprung-Russell (HR) diagram both for the same metallicity and over a range of metallicities. These uncertainties are compounded by the fact that errors in spectroscopically-determined parameters (such as [Fe/H]) grow as the stellar absorption lines become more rotationally broadened. Knowing the stellar mass is important because of the very different rotation evolution followed by stars of different masses. Stars more massive than $\sim$1.3~\msun\ retain the majority of their birth AM throughout their main sequence (MS) lifetimes, resulting in both large average rotational velocities and a large dispersion in rotational velocities (e.g., \citealt{Kraft:1967jh}, \citealt{Royer:2007dj}). Even within this intermediate mass regime, different rotational velocities are found. \cite{Gray:1982bc} inferred from the distribution of rotation with mass that the stellar AM ($J$) followed a power law with mass ($M \propto J^{5/3}$), with a break to a steeper power law below 1.6~\msun. \cite{1997PASP..109..759W} demonstrated that the lower specific AM seen at 1.3--1.6~\msun\ was mostly imparted during the pre-MS stages, but also noted that some slow loss of AM might occur during the MS. Conversely, stars less massive than $\sim1.3$~\msun\ rapidly spin down, eventually erasing all information on their birth AM and allowing the field of gyrochronology to use measured rotation as a proxy for age on the MS (e.g., \citealt{Barnes:2003ga}, \citealt{Barnes:2007fz}, \citealt{Chaname:2012jz}). Given the larger AM seen in more massive MS stars, it seems logical to assume that the field population of fast RG stars are simply the most massive stars. One way of testing this assumption would be to measure the rotation of subgiant stars crossing the Hertzsprung gap, which would provide a direct measure of the spin-down from the MS to the base of the RGB. However, these stars are relatively rare. To study a population of RGs with well-constrained masses, open cluster stars are needed. To date, little attention has been paid to the distribution of \vsini\ of RGs in open clusters. Therefore, in this study over 400 red giant candidate members of eleven open clusters were chosen to probe the rotation distribution in open clusters (Section \ref{sec:samples}). High resolution spectra were obtained for these candidates (Section \ref{sec:obs}), from which heliocentric radial velocities (\vhelio) and \vsini\ are measured (Section \ref{sec:measure}). The \vhelio\ distributions are used to define the cluster's \vhelio\ and identify likely members (Section \ref{sec:vhelio}). The rotation distributions are analyzed with respect to both stellar masses and current evolutionary stages (Section \ref{sec:vrot_results}), and potential explanations for the fastest rotators are explored (Section \ref{sec:sources}). Finally, the main results and conclusions of the study are summarized (Section \ref{sec:summary}). | \label{sec:summary} We surveyed the rotation distribution of RGs in eleven open clusters to search for rapid rotators in an effort to better understand the outliers of this otherwise generally slow-rotating class of stars. Rapid and moderately fast rotators appear to be as common in clusters as in the field population, but the frequency of moderate rotators in Pismis~18 is surprisingly large. Separating the stars into mass bins with different average rotation on the main sequence, we find that the more massive stars ($M\gtrsim1.6$~\msun) have a larger population of enhanced rotators, and the rotation distribution is much broader. However, when converting to AM and restricting the comparison to stars with similar sizes, the differences, while still present, are not statistically significant. Nevertheless, there are clear outliers to the general AM distributions among the higher mass stars. These data provide constraints on models of stellar rotation of intermediate mass stars, which appear to spin down very rapidly during the subgiant evolution phase. Stars in the subgiant phase are quite rare, and their radii do not change drastically enough to account for the amount of spin-down necessary to explain the slow rotation seen in the earliest RG phase. If some of the AM is sequestered in the stars instead of lost by magnetized winds, evolution models can use the rotation distribution of RGs to test how that AM may resurface. Measuring additional rotational velocities or rotation periods of intermediate mass stars crossing the Hertzsprung Gap remains a difficult but important piece of the puzzle. Distinguishing between internal and external sources of enhanced surface rotation may come from disentangling the ambiguous stage of evolution that may either be first dredge-up or the red clump. AM dredge-up may be at work in the former case, while the latter may require an external AM origin. Testing these two possibilities could be accomplished with asteroseismic analysis of a rapidly rotating RG to distinguish whether the He core is inert or is an active region of nucleosynthesis. | 14 | 4 | 1404.4358 |
1404 | 1404.4996_arXiv.txt | Recent 3D-simulations have shown that the turbulent kinetic flux (TKF) is significant. We discuss the effects of TKF on the size of convection zone and find that the TKF may help to solve the solar abundance problem. The solar abundance problem is that, with new abundances, the solar convection zone depth, sound speed in the radiative interior, the helium abundance and density in the convective envelope are not in agreement with helioseismic inversions. We have done Monte Carlo simulations on solar convective envelope models with different profile of TKF to test the effects. The solar abundance problem is revealed in the standard solar convective envelope model with AGSS09 composition, which shows significant differences ($\rm{\sim 10 \%}$) on density from the helioseicmic inversions, but the differences in the model with old composition GN93 is small ($\rm{\sim 0.5 \%}$). In the testing models with different imposed TKF, it is found that the density profile is sensitive to the value of TKF at the base of convective envelope and insensitive to the structure of TKF in the convection zone. Required value of turbulent kinetic luminosity at the base is about $\rm{-13\%\sim-19\%L_{\odot}}$. Comparing with the 3D-simulations, this value is plausible. This study is for the solar convective envelope only. The evolutionary solar models with TKF are required for investigating its effects on the solar interior structure below the convection zone and the whole solar abundance problem, but the profile of TKF in the overshoot region is needed. | \label{sec1} Recent photospheric analysis (e.g., \citet{AGS05,AGSS09,caf11}) have indicated that the solar photospheric metallicity is significantly lower than the old values \citep{GN93,GS98}. This arises the solar abundance problem that standard solar models with the revised composition show serious deviations from the helioseismic inversions, i.e., the depth of the convection zone (CZ) $\rm{r_{bc}}$, the surface helium abundance $\rm{Y}$ and the sound speed and density in the solar interior \citep{ba04,bah05,bah06,yb07}. Many models are proposed to modify the solar model, including enhanced diffusion \citep{guz05}, the accretion model \citep{guz05,guz10,ser11}, particles with an axion-like interaction \citep{Vincent13}, etc.. However, none has succeeded in solving the problem. The accretion model have shown improvements on the solar model, but the inconsistence remains since $\rm{r_{bc}}$ and $\rm{Y}$ can not fit the helioseismic restrictions simultaneously \citep{guz05,guz10,ser11}. It is expected that the opacity at the base of the convection zone (BCZ) could be adjusted upward, since the tests of enlarged opacity reduce the discrepancies between model and helioseismic inversions \citep{ba04,bah05,chr09}. The turbulent kinetic flux (TKF) is usually ignored in modeling star because it is thought to be small. Another reason is that there is no widely accepted nonlocal convection theory to describe the TKF. However, 3D-simulations on stellar convection envelopes have shown that TKF is not ignorable in some cases \citep{tian09,hotta14}. It should be noticed that the TKF at BCZ is negative because the turbulent convection transport kinetic energy from the CZ to the radiative region. A negative TKF requires radiative flux plus convective flux being larger, thus leads to a deeper convective boundary. This could help to improve the solar model. Based on this ideal, we test the TKF in solar convective envelope (CE) models. | \label{sec5} In this letter, we discussed the effects of the turbulent kinetic flux (TKF) on the size of convection zone (CZ) and test the effects of TKF on the solar convective envelope (CE) models. The main conclusions are as follows: (i) The presence of TKF modifies the convective criterion and makes convective boundaries shift downward, thus the convective core becomes smaller and the CE becomes larger. (ii) The solar abundance problem is revealed in the solar CE models. The standard solar CE model with revised composition \citep{AGSS09} shows significant difference ($\rm{\sim 10\%}$) on density from the helioseismic inversions. This makes it be impossible to obtain a solar model with AGSS09 composition fitting all helioseismic restrictions if the standard stellar structure equations and standard input physics are used in the solar CE. (iii) Taking into account TKF could improve the solar CE model. The density structure of the solar CE is sensitive to the value of the TKF at the BCZ and insensitive to its profile in the CZ. Required turbulent kinetic luminosity at the BCZ is $\rm{-13\% L_{\odot} < L_{K,bc} < -19\% L_{\odot} }$ taking into account the uncertainties of $\rm{Y}$ and $\rm{r_{bc}}$. In this paper, we have done a limited test with the TKF on the solar abundance problem, i.e., testing required TKF to construct the convective envelope of the solar model (with AGSS09 composition) fitting all helioseismic restrictions. However, we can't show the effect of the TKF in complete solar evolutionary models. The TKF profile below the BCZ is required to do that. However, there is no simulation showing the profile. The standard model with AGSS09 abundance show lower helium abundance and shallower BCZ comparing with the helioseismic inversions. \citet{guz05} have shown that the sound speed differences below the BCZ can not be removed in some non-standard models even with the correct BCZ. We think that taking into account some important aspects of convection (the overshoot mixing and turbulent kinetic flux) could help to solve the solar abundance problem. The incomplete mixing caused by the convective overshoot \citep{zha13} partially compensates the settling thus enlarge the helium abundance. And the incomplete mixing below the BCZ is favored by the sound speed when the BCZ is in the correct location \citep{brun99,zha12,zha13}. A fault of the incomplete mixing is to lead to a low Z when the surface $\rm{Z/X}$ is fixed, thus the BCZ becomes shallow \citep{brun99,zha12}. However, the TKF could compensate it. The required $\rm{L_{K,bc}}$ seems to be too high that is comparable with the total luminosity. According to \citet{tian09} and \citet{hotta14}, that is plausible. Xiong's turbulent convection model gives $\rm{L_{K,bc}} \sim -1\% L_{\odot}$ \citep{xiong01}. However, \citet{tian09} have shown that the gradient type model of the TKF in Xiong's model is too imprecise to be acceptable. A serious problem that the stellar models could be significantly affected is arising if the TKF is in true far away from ignorable. The possibly effects may be in H/He main-sequence stars with convective core and RGB/AGB stars with convective dredge up, since the variation of the boundary of the convective core/envelope changes the profile of chemical abundance in stellar interior. | 14 | 4 | 1404.4996 |
1404 | 1404.3863_arXiv.txt | {Active Galactic Nuclei (AGN) are the most luminous persistent objects in the universe. The X-ray domain is particularly important as the X-ray flux represents a significant fraction of the bolometric emission from such objects and probes the innermost regions of accretion disks, where most of this power is generated. An excess of X-ray emission below $\sim$ 2 keV, called soft-excess, is very common in Type 1 AGN spectra. The origin of this feature remains debated. Originally modeled with a blackbody, there are now several possibilities to model the soft-excess, including warm Comptonization and blurred ionized reflection. In this paper, we test ionized-reflection models on Mrk 509, a bright Seyfert 1 galaxy for which we have a unique data set, in order to determine whether it can be responsible for the strong soft-excess. We use ten simultaneous XMM-Newton and INTEGRAL observations performed every four days. We present here the results of the spectral analysis, the evolution of the parameters and the variability properties of the X-ray emission. The application of blurred ionized-reflection models leads to a very strong reflection and an extreme geometry, but fails to reproduce the broad-band spectrum of Mrk 509. Two different scenarios for blurred ionized reflection are discussed: stable geometry and lamp-post configuration. In both cases we find that the model parameters do not follow the expected relations, indicating that the model is fine-tuned to fit the data without physical justification. A large, slow variation of the soft-excess without counterpart in the hard X-rays could be explained by a change in ionization of the reflector. However, such a change does not naturally follow from the assumed geometrical configuration. Warm Comptonization remains the most probable origin of the soft-excess in this object. Nevertheless, it is possible that both ionized reflection and warm Comptonization mechanisms can explain the soft-excess in all objects, one dominating the other one, depending on the physical conditions of the disk and the corona.} | In non-obscured Active Galactic Nuclei (AGN), most of the radiation is emitted in the optical-UV and the X-ray energy bands. In the optical-UV band, the emission is characterized by the ``big blue bump", present from about 10 nm to 0.3 $\mu$m \citep{Sanders1989,Bregman1990,Zhou1997}. This emission is thought to come from an optically thick accretion disk \citep{ShakuraSunyaev1973}. The X-ray spectrum of Seyfert galaxies is typically characterized by a power-law continuum with reflection features, absorption and often an excess in the soft X-rays \citep{Halpern1984,Turner1989}. The power-law emission is believed to be due to Comptonization of the UV photons coming from the disk by the energetic electrons in a corona surrounding the disk \citep{Blandford1990,Zdziarski1995,Zdziarski1996,Krolik1999}. As a high-energy cut-off is detected in a range from about 80 to 300 keV in about 50\% of Seyfert galaxies, a thermal distribution is preferred to a non-thermal one \citep{Gondek1996,Matt2001,Perola2002}. The strongest signatures of the reflection component are a ``reflection hump" around 30 keV and an iron FeK$\alpha$ fluorescence line in between 6 and 7 keV depending on the iron ionization state. Reflection is associated with the reprocessing of the primary continuum by material either close to the central black hole, in the accretion disk \citep{George1991,Matt1991} or more distant, for example in a torus \citep{Antonucci1993,Jaffe2004,Meisenheimer2007,Raban2009}. It can also be produced either in the narrow or broad line regions \citep{Ponti2013}. Absorption from material either in the vicinity of AGN or in the host galaxy is generally observed in the X-ray spectra of Seyfert 1 and 2 galaxies. If the absorbing material is often photoionized, it is referred to as ``warm absorber" \citep{George1998}. More than 50\% of Seyfert 1 galaxies show the presence of a soft X-ray excess called soft-excess \citep{Halpern1984,Turner1989}, a soft X-ray emission below $\sim$ 2 keV in excess of the extrapolation of the hard X-ray continuum. \cite{Piconcelli2005} and \cite{Bianchi2009} even found that the fraction of AGN with soft-excess reaches about 100\%. The discovery of this component was made thanks to the HEAO-I \citep{Singh1985} and EXOSAT \citep{Arnaud1985} missions in the 80's, but its nature is still uncertain. It was first thought to arise from the hottest part of the accretion disk \citep[e.g.][]{Arnaud1985,Pounds1986}, but this hypothesis was invalidated by the facts that the temperature of the soft-excess (0.1-0.2 keV) is much too high to be explained by the standard accretion disk model around a supermassive black hole, and that it does not vary, as expected, with the mass of the black hole \citep{GierlinskiDone2004}. Another possible explanation is ``warm" Comptonization: up-scattering of seed disk photons in a Comptonizing medium which has a temperature of about 1 keV (e.g. in NGC 5548 -- \citealt{Magdziarz1998}; RE J1034+396 -- \citealt{Middleton2009}; RX J0136.9--3510 -- \citealt{Jin2009}; Ark 120 -- \citealt{Matt2014}; and 1H 0419--577 -- \citealt{DiGesu2014}). This Comptonization model is supported by the fact that strong similarities have been found in the optical-UV and soft X-ray variability, suggesting a correlation between these emissions, in agreement with inverse Compton processes \citep{Edelson1996,Mehdipour2011}. \cite{Walter1993} used a sample of 58 Seyfert 1s observed by ROSAT and IUE and found a spectral shape correlation (i.e. an amplitude correlation) indicating that the soft-excess could be the high-energy tail of the "big blue bump" component observed in UV, as objects with a large UV bump do also show a strong soft-excess. \cite{Edelson1996} found a correlation between variabilities in X-rays, UV and optical bands in NGC 4151. \cite{Marshall1997} studied NGC 5548 in the extreme UV band using EUVE and noticed that the EUV and UV/optical variations of the light curves are simultaneous. They also found that the shape of the EUVE spectrum is consistent with that in UV and soft X-rays. However, the warm Comptonization model still does not explain why the shape of the soft-excess does not appear to vary with the black-hole mass \citep{GierlinskiDone2004}. An alternative explanation is that the soft-excess is linked to atomic processes. The soft-excess could be the signature of strong, relativistically smeared, partially ionized absorption in a wind from the inner disk. \cite{GierlinskiDone2004} applied this model on PG 1211+143. If a totally covering absorber is assumed, the problem of this model is that it needs quite extreme values of the model parameters to account for the observed smooth soft-excesses, in particular a very large smearing velocity, which is not attainable in models of radiatively driven accretion disk winds with typical physical parameters \citep{Schurch2007,Schurch2009}. This problem does not apply anymore in the case of a partial-covering scenario. In Mrk 766, for example, the principal components analysis and spectral variability of a long XMM-Newton observation can be explained by ionized absorption partially covering the continuum source \citep{Miller2007}. \cite{Turner2009} present a review on X-ray absorption and reflection in AGN, showing that partial-covering absorption can explain spectral curvature and variability at low energy. Another interpretation is that the soft-excess is the result of ionized reflection in a relativistic disk, which blurs all emission lines. An ionized-reflection model, calculated for an optically-thick atmosphere of constant density illuminated by radiation of a power-law (called \textit{reflionx} in XSPEC; \citealt{RossFabian2005}), has been successfully applied by \cite{Crummy2006} on 22 Type-1 PG quasars and 12 Seyfert 1 galaxies and by \cite{Zoghbi2008} on Mrk 478 and EXO 1346.2+2645. This model has also been applied in MCG-6-30-15 by \cite{Vaughan2004} and NGC 4051 by \cite{Ponti2006}, explaining the spectral shape as well as the variability. The spectral and timing analysis of 1H 0707-495 \citep{Fabian2009} provides strong evidence of emission from matter close to a rapidly spinning black hole. This object shows the presence of a broad iron K line whose width and shape is a signature of strong gravity and spin of the black hole. Thanks to the high iron abundance, the iron L line is also detectable and a lag of about 30 seconds with the direct X-ray continuum could be measured. This lag is an evidence of reverberation processes. X-ray reverberation time delays have also been detected in MCG--5--23--16 and NGC 7314 by \cite{Zoghbi2013} using the iron K$\alpha$ emission lines. \cite{Kara2013} also found iron K lags in Ark 564 and Mrk 335. Soft X-ray reverberation lags have been found in ESO 113-G010 by \cite{Cackett2013} and PG 1211+143 by \cite{DeMarco2011}. Overall, soft lags have been found in more than 15 sources \citep{DeMarco2013}. Blurred ionized reflection has also been tested, as well as double Comptonization and ionized absorption by a high velocity material, in Mrk 509 and Mrk 841, using average broad-band Suzaku data \citep{Cerruti2011}. This model seems to correctly describe Mrk 509 soft-excess, but underestimates the broad iron emission line. In this paper, we focus on testing blurred ionized-reflection models in an object for which we have a unique data set: Mrk 509. We want to check whether ionized reflection could be a viable alternative explanation to the warm Comptonization to explain the soft-excess. In order to test these ionized-reflection models, we use the Mrk 509 XMM-Newton and INTEGRAL campaign and its 10 simultaneous observations \citep{Kaastra2011}. We first study the variability at different time scales and then we test blurred ionized-reflection models on the ten observations in order to investigate the evolution of the \textit{reflionx} parameters. The paper is organized as follows. In Sect. \ref{2}, we first introduce the properties of Mrk 509 and the XMM-Newton/INTEGRAL campaign of observation that we use. In Sect. \ref{3}, we show the results of the study on variability at different time scales. Section \ref{4} presents the ionized-reflection models that we use to fit our data, as well as the results on the average spectrum and for each observation. In Sect. \ref{5}, we show the evolution of \textit{reflionx} parameters, taking into account two cases: the case of a stable geometry with a constant reflection factor and the case of a varying reflection factor. All these results are discussed in Sect. \ref{6} and we state the conclusions in Sect. \ref{7}. | \label{7} The nature of the soft-excess in AGN is still uncertain. While it can be fitted by both ionized reflection \citep{Crummy2006,Zoghbi2008,Fabian2009,Ponti2010} and warm Comptonization \citep{Magdziarz1998,Middleton2009,Jin2009,Mehdipour2011,Petrucci2012} in most, if not all, objects, some detailed features observed in individual objects point either to the former -- lags in the soft-excess \citep{Cackett2013,DeMarco2011,DeMarco2013}; in iron K and L lines \citep{Fabian2009,Zoghbi2013,Kara2013} -- or to the latter -- correlation with UV or variability spectrum \citep{Walter1993,Edelson1996,Marshall1997,Mehdipour2011,Done2012}. Because blurring can make ionized reflection look very featureless, the distinction between the two models at low energies is difficult and leads to a confusing situation about the origin of the soft-excess. In this work we applied the blurred ionized-reflection model in Mrk 509, a bright Seyfert 1 galaxy for which we have a unique data set and which shows the presence of a strong soft-excess. The ionized-reflection model has some difficulty in fitting the broad-band spectrum of Mrk 509, even assuming a very strong reflection and an extreme geometry. We first made the hypothesis of a stable geometry, but this configuration leads to a non-physical anti-correlation between the ionization parameter and the \textit{reflionx} normalization and cannot explain the strong variability of the soft-excess on long time scales. The soft-excess variability cannot be explained by a varying geometry such as a lamp-post configuration either. Furthermore, even if light-bending effects can induce a high reflection factor value, we cannot find the expected correlation between \textit{R*} and $\xi$. The \textit{reflionx} model fine-tunes its parameters in order to fit the data, introducing non-physical relations between parameters and preventing expected relations to appear. Ionized reflection is then unable to explain the origin of the soft-excess in Mrk 509. In \cite{Mehdipour2011} and in \cite{Petrucci2012}, the soft-excess of Mrk 509 was attributed to warm Comptonization based on some observed relationships, in particular the correlation between the UV and the soft X-ray fluxes. Under this hypothesis, the excess variability in the soft X-ray flux on long time scales can be explained by changes in the accretion rate (and hence in the seed photon flux). This model of warm Comptonization remains therefore the most probable explanation of the soft-excess in this object. Ionized reflection is expected as soon as a strong X-ray source irradiates very nearby cold-to-warm matter. Therefore, it is well possible that both mechanisms are working in all objects, but the dominance of one over the other depends on the physical conditions of the disk and of the corona. The advent of sensitive telescopes in the hard X-rays like NuSTAR and in the future ASTRO-H shall provide very useful constraints on the origin of the soft-excess. One physical parameter that can be expected to determine the existence of blurred ionized reflection is the Eddington ratio. Low Eddington ratios may indeed imply a transition from a standard accretion disk to an advection-dominated accretion flow \citep{Ichimaru1977,Rees1982,Narayan1994,Narayan1995a, Narayan1995b,Abramowicz1995}, effectively truncating the accretion disk and removing cold material from the inner parts where relativistic effects are most important. This is further supported by the fact that the best evidences for blurred ionized reflection, including reverberation lags in the soft-excess \citep{Cackett2013,DeMarco2011} or iron L and K emission \citep{Fabian2009,Zoghbi2013,Kara2013}, are found in Narrow-Line Seyfert 1 objects (for instance 1H 0707--495 and PG 1211+143), which are thought to have very high accretion rates. The moderate Eddington ratio of 0.3 in Mrk 509 may therefore be too low for blurred ionized reflection to become the dominant source of soft-excess in this object. However, lags have also been detected in many Seyfert galaxies without high accretion rates \citep{DeMarco2013}. The ten simultaneous XMM-Newton and INTEGRAL observations of Mrk 509 represent a unique data set which brings unprecedented constraints on the models. Many models are able to reproduce the soft-excess, because it is essentially featureless. Monitoring campaigns allow to study emission models along the time dimension, and can provide very useful constraints to determine the origin of components like the soft-excess. Additional monitoring campaigns on other Seyfert 1 objects, and in particular narrow-line Seyfert 1 objects, where the evidence for ionized reflection is the strongest, would be very useful to understand whether this component can really have different origins in different objects. | 14 | 4 | 1404.3863 |
1404 | 1404.4444_arXiv.txt | We present 2MASS $JHK_{\rm s}$ photometry for 913 star clusters and candidates in the field of M31, which are selected from the latest Revised Bologna Catalog of M31 globular clusters (GCs) and candidates. The photometric measurements in this paper supplement this catalog, and provide a most comprehensive and homogeneous photometric catalog for M31 GCs in the $JHK_{\rm s}$ bandpasses. In general, our photometry is consistent with previous measurements. The globular cluster luminosity function (GCLF) peaks for the confirmed GCs derived by fitting a $t_5$ distribution using maximum likelihood method are: $J_0 = 15.348_{-0.208}^{+0.206}$, $H_0 = 14.703_{-0.180}^{+0.176}$, and ${K_{\rm s}}_0 = 14.534_{-0.146}^{+0.142}$, all of which agree well with previous studies. The GCLFs are different between metal-rich (MR) and metal-poor (MP), inner and outer subpopulations, as that MP clusters are fainter than their MR counterparts, and the inner clusters are brighter than the outer ones, which confirm previous results. The NIR colors of the GC candidates are on average redder than those of the confirmed GCs, which lead to an obscure bimodal distribution of the color indices. The relation of $(V-K_{\rm s})_0$ and metallicity shows a notable departure from linearity, with a shallower slope towards the redder end. The color-magnitude diagram (CMD) and color-color diagram show that many GC candidates are located out of the evolutionary tracks, suggesting that some of them may be false M31 GC candidates. The CMD also shows that the initial mass function of M31 GCs covers a large range, and the majority of the clusters have initial masses between $10^3$ and $10^6$ $M_{\odot}$. | \label{intro.sec} Globular clusters (GCs) provide a unique laboratory for investigating the formation and evolution of their host galaxies. The form of the mass spectrum for GCs is nearly identical to the mass function of the parent molecular cloud cores \citep{mp96}. The brightness distribution of GCs, known as globular cluster luminosity function (GCLF), can be used to constrain possibilities for GC formation and destruction \citep[see][and references therein]{nantais06}. By virtue of the natural advantage of being located at a reasonable distance, nearby galaxies (especially M31) offer us an ideal environment for detailed studies of cluster systems. A large number of GCs have been identified in M31 since \citet{hubble32}, and the latest Revised Bologna Catalog of M31 GCs and candidates \citep[hereafter RBC V.5,][]{gall04,gall06,gall09}, which is a compilation of photometry and identifications from many previous catalogs, published 625 confirmed GCs and 331 GC candidates. Based on the growing number and updated sample of M31 GCs and candidates, many works have probe the GCLF in M31 in detail. \citet{crampton85} found that the mean luminosity of GCs in M31 decreases markedly with increasing galactocentric distance, indicating that clusters with larger projected distances are fainter than those closer to the nucleus \citep{gnedin97}. \citet{og97} obtained distance moduli to M31 and M87 from GCLF by applying the corrections for dynamical evolution, and found surprising consistency of the predicted and observed distances, which confirmed the GCLF as a distance indicator. These authors also concluded that the mass functions of GCs in the Milky Way (MW), M31, and M87 were universal at the birth of these systems, although spanning a wide range of masses. \citet{bhb01} measured the LF for M31 GCs, and found that inner clusters have a GCLF peak brighter than the outer ones, while the metal-rich (MR) clusters are brighter than their metal-poor (MP) counterparts. The variation in the M31 GCLF seems to be due to various factors: metallicity, age, cluster initial mass function (IMF), and dynamical destruction \citep{bhb01}. Recently, \citet{nantais06} compared the GCLFs of the MW, M31, and the Sculptor Group spiral galaxies, and found that the GCLF of the MW is consistent with that of M31. Near-infrared (NIR) colors can help to distinguish among star formation histories and IMFs \citep{Barton03}. The Two Micron All Sky Survey (2MASS), performed between 1997 June and 2001 February, covers 99.998\% of the celestial sphere in the $J$ (1.25 $\mu$m), $H$ (1.65 $\mu$m), and $K_{\rm s}$ (2.16 $\mu$m) bandpasses \citep{Skrutskie06}. These observations were conducted with two 1.3 m diameter telescopes located at Arizona and Chile. The 2MASS All-Sky Data Release contains the observation of M31 with an integration time of 7.8 s for each exposure, while a new 2.8 deg$^2$ NIR survey from the 2MASS 6X program across the extent of the optical disk of M31, with an exposure time of 6 times the normal exposure, provides a clearer view of the galaxy center. \citet{gall04} identified 693 known and candidate GCs in M31 using the 2MASS database, and derived their 2MASS $JHK_{\rm s}$ magnitudes. After adding the mean difference between the 2MASS photometry and previous NIR photometry, the newly assembled NIR dataset were implemented into a revised version of Bologna Catalog. These authors also showed that the $V-K_{\rm s}$ color provides a powerful tool to discriminate between M31 clusters and background galaxies. \citet{santos13} presented 2MASS photometry and color for a sample of Local Group clusters younger than $\sim$100 Myr, and found that the embedded clusters, which are heavily obscured by dust, generally have a redder $H-K_{\rm s}$ color than older ones, from which gas and dust have already been ejected. These authors also concluded that the brightest clusters can be split into young and old subsamples from $H-K_{\rm s}$ color. Considering that a more extended and homogenous photometry in the NIR is important for the studies on M31 star clusters, we would carry out new photometry for them using 2MASS images. In this paper, we provide NIR photometry for a set of star clusters in M31 using images from 2MASS. This paper is organized as follows. In Section 2 we present the sample selection, $JHK_{\rm s}$ photometry, and comparisons with previous measurements. A discussion on the properties of the sample clusters will be given in Section 3, including the GCLF and color distributions. Finally, we will summarize our results in Section 4. | \label{discussion.sec} We combined the photometry results of $JHK_{\rm s}$ bands newly derived here with those derived by \citet{ma12b} for 10 GCs in the M31 halo to construct a more comprehensive sample of 923 clusters in M31 to discuss the properties for them. The reddening values for the 10 GCs were derived from \citet{ma12b}, while the metallicities from \citet{cald11} and \citet{kang12} were used for them, respectively. The $V$-band magnitudes of the sample clusters were derived from RBC V.5 for the following analysis about the color distribution. All these magnitudes have been extinction corrected. \subsection{Luminosity Function} \label{LF.sec} To obtain the M31 GCLF peaks, we used least square and maximum likelihood methods to fit a Gaussian and Student's t distribution on the dereddened $JHK_{\rm s}$ data. The Gaussian distribution is given as \begin{equation} f(x)={\frac{1}{\sqrt{2\pi}\sigma_{G}}} {\rm exp}(-{\frac{(x-\mu_{G})^2}{2\sigma_{G}^2}}) \end{equation} The Student's t distribution is defined by \begin{equation} f(x)={\frac{\Gamma((n+1)/2)}{\sqrt{\pi~n}\Gamma(n/2)\sigma_{t}}} (1+{\frac{(x-\mu_{t})^2}{n\sigma_{t}^2}})^{-(n+1)/2} \end{equation} where n is the degree of freedom (DOF). \citet{secker92} and \citet{secker93} reported that the Student's t distribution with DOF of 5 (hereafter ``$t_5$''), which presents a power-law fall-off in its wings, is more robust than the Gaussian in describing the outlying data points and estimating the GCLF peak. The ``$t_5$'' function is evaluated as \begin{equation} f(x)={\frac{8}{3\sqrt{5}\pi\sigma_{t}}} (1+{\frac{(x-\mu_{t})^2}{5\sigma_{t}^2}})^{-3} \end{equation} Some other distributions \citep{baum95, Larsen01} were also investigated for the GCLF. \citet{baum95} found that a composite of two exponentials is a better fit over the Gaussian or Student's t distribution to the combined LF of the Galactic and M31 GCs, since those two distributions fail to deal with the asymmetry and with the sharpness of the peak of the histogram. Figure 9 displays the luminosity histograms and the best fitting lines for the sample clusters. The top panels show the fitting with least square technique, while the bottom panels show the fitting with maximum likelihood method. The black lines show the fitting with a $t_5$ distribution, while the red lines show the fitting with a Gaussian distribution. The maximum likelihood method is used to estimate the most probable parameter values from the sample data \citep{secker92}, which would not suffer from the effect of binning. As shown in Figure 9, the fitting with a $t_5$ distribution shows a more extended wing than that with a Gaussian distribution \citep{secker92}. Table 3 lists the LF parameters for the sample clusters, the confirmed GCs, and those confirmed GCs with metallicities available from \citet{kang12}. The LF peaks for confirmed GCs are much brighter than those for all clusters, indicating that a clean sample of GCs is critical to obtain accurate LF parameters. The GCLF peaks derived by \citet{bhb01} are $J_0 = 15.26$ and $K_0 = 14.45$, by fitting a $t_5$ distribution using the MAXIMUM program written by J. Secker \citep[see][for details]{secker92}. \citet{nantais06} derived similar results for the $JHK_{\rm s}$ bands using same methods: $J_0 = 15.31$, $H_0 = 14.76$, and ${K_{\rm s}}_0 = 14.51$, and these authors concluded that the little fainter peaks than those from \citet{bhb01} could be real, due to a larger and possibly deeper sample. In this paper, the LF peaks for the confirmed GCs, derived by fitting a $t_5$ distribution using maximum likelihood method, are: $J_0 = 15.348_{-0.208}^{+0.206}$, $H_0 = 14.703_{-0.180}^{+0.176}$, and ${K_{\rm s}}_0 = 14.534_{-0.146}^{+0.142}$, all of which agree well with previous results. The $\sigma$ parameters derived here are slightly larger than previous studies. \begin{figure}[!htb] \figurenum{9} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig9.ps}}} \vspace{0.1cm} \caption{The luminosity histograms and the best fitting models for the sample clusters. The top panels show the fitting with least square technique, while the bottom panels show the fitting with maximum likelihood method. The black and red lines show the fitting with a $t_5$ and Gaussian distribution, respectively.} \label{fig:fig9} \end{figure} It has long been known that the peak of the GCLF is nearly constant in different galaxies, providing a standard indicator for the cosmological distance measurement \citep{racine68, hanes77, ferrarese00}. However, several studies \citep{crampton85, gnedin97, bhb01} presented that GCLF varies between MR and MP, inner and outer subsamples within a galaxy, and GCLF peak becomes fainter as the local density of galaxies increases \citep{bt96}. \citet{Larsen01} found that the $V$-band turnover of the blue GCs is brighter than that of the red ones by about 0.3 mag on the average, with a study of GCs in 17 nearby early-type galaxies. \citet{bhb01} found that MR clusters are brighter than MP ones in M31, and inner clusters are brighter on average than outer clusters, indicating that the luminosity function is different among these subpopulations. \citet{Goudfrooij04} confirmed that the GC system in the early-type galaxy NGC 1316, which is an intermediate-age merger remnant, can be divided into a blue GC subpopulation, consistent with a Gaussian LF, and a red GC component with a power law LF. These authors also found that the LF of the inner half of the MR population differs significantly from that of the outer half. The difference in GCLF between dwarf and giant ellipticals has been studied by many authors \citep{harris91, durrell96, strader06, jordan06, ml07}, however, whether there is a trend of the GCLF peak with the galaxy luminosity is still under debate. To investigate the effects of metallicity and galactocentric distance on the LF, we divided the confirmed GCs into four subsamples as \citet{bhb01} did. The MR and MP subpopulations are divided at [Fe/H] $=-1.1$, using metallicities from \citet{kang12}, while the inner and outer subsamples are divided at $R_{\rm gc}=10$ kpc. Figure 10 displays the $J_0$ luminosity histograms for the four subsamples, together with the best-fit $t_5$ distribution using the maximum likelihood method. Table 4 lists the GCLF fit results in the $JHK_{\rm s}$ bands for the four subsamples. It is evident that the MP clusters are fainter than their MR counterparts, and the inner clusters are brighter than the outer ones. \citet{gnedin97} reported that the destruction of the inner faint, low-density clusters due to strong tidal shocks may lead to the difference between the LF of the inner and outer populations. \citet{kh97} asserted that the difference in the GCLF of the inner and outer halo populations is due to dynamical evolution and/or a dependence of GCLF shape on the environment. \citet{og97} found that the predicted differences between the peaks of the inner and outer cluster populations, due to the tidal shocks and dynamical friction, agree quantitatively with the observed differences within the errors in the MW, M31, and M87. However, \citet{bhb01} concluded that metallicity, age, and cluster IMF may also be important for the variation in the M31 GCLF except dynamical destruction. \begin{figure}[!htb] \figurenum{10} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig10.ps}}} \caption{The $J_0$ luminosity histograms and the best fitting models for four groups of the confirmed GCs, divided by the galactocentric distances and metallicities from \citet{kang12}. The black lines show the fitting with a $t_5$ distribution using the maximum likelihood method.} \label{fig:fig10} \end{figure} \subsection{Color Distribution} \label{color.sec} Figure 11 displays the NIR color (including $(V-K_{\rm s})_0$ ) distribution of the whole sample clusters (solid-line histogram) and the confirmed GCs (filled histogram) with magnitude uncertainty $< 0.5$ mag. There are some bumps located in the redder wings of the $(J-K_{\rm s})_0$, $(H-K_{\rm s})_0$, and $(V-K_{\rm s})_0$ colors, which are mainly caused by GC candidates. The GC systems of many galaxies reveal bimodal optical color distributions \citep{fbh97, kundu99, gk99, Larsen01, rejkuba01}. \citet{bh00} have detected the bimodal distribution of the GC colors in M31 using $(U-V)_0$, $(U-R)_0$, and $(V-K)_0$ with the KMM statistics \citep{mb88, ashman94}. To check whether the NIR color distributions are bimodal, the KMM algorithm was also performed here. The homoscedastic fitting, with same variances for both groups, was assumed. Table 5 lists the parameters returned by the KMM algorithm. Column (2) and (3) give the estimated mean value and covariance assuming the whole sample as one group. Column (4) and (5) give the estimated mean value for each group, assuming the whole sample as two groups, and Column (6) gives the common covariance for them. Column (7) and (8) give the number points assigned to each group. Column (9) gives the $p$ value, which is an estimate of the improvement of the two-group fit over a one-group fit, and is interpreted as a rejection of the single Gaussian model at a confidence level of $1-p$. The purely NIR color distributions show bimodality at the $\sim$100 per cent confidence level, however, the size of the redder subsample is very small, and this subsample is composed mainly of the GC candidates, as shown in Figure 11. Although Table 5 presents the mean values and number points for the two groups of $(V-K_{\rm s})_0$ color, these values vary with the initial set, and the $p$ value is unavailable since the sample is not convergent. Some previous studies of purely NIR and optical-NIR GC colors in different galaxies have shown that in some cases the optical color distributions are clearly bimodal, however, the purely NIR and optical-NIR color distributions are not, or they display ``differing bimodalities'' \citep{blakeslee12} with those of the optical colors alone \citep[see][and references therein]{cantiello}. \begin{figure}[!htb] \figurenum{11} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig11.ps}}} \caption{The color distribution of the sample clusters.} \label{fig:fig11} \end{figure} Figure 12 shows the relation between metallicity and intrinsic colors for the sample clusters with magnitude uncertainty $< 0.5$ mag. Metallicities in the top panels are from \citet{kang12}, while metallicities in the bottom panels are from \citet{cald11}. The open circles represent the confirmed GCs, while the crosses represent the rest clusters, including candidate GCs, controversial objects, and extended clusters. The $(V-K_{\rm s})_0$ color index, which is often used as a metallicity indicator, shows a clearer correlation with [Fe/H] than other NIR colors, indicating that these NIR intrinsic colors are less sensitive to metallicity \citep{bh00}. \citet{nantais06} also reported that stellar spectra in the NIR is much less dependent on metallicity than in the optical. The relation of $(V-K_{\rm s})_0$ and metallicity shows a notable departure from linearity, with a shallower slope in the redder part. The nonlinear correlation of color with metallicity, which are mainly driven by the horizontal-branch stars, can produce a bimodal color distribution with unimodal metallicity for a group of old clusters \citep{Yoon06}. We also use the KMM algorithm to investigate the bimodality of the metallicity distributions for M31 GCs. Table 6 lists the parameters returned by the KMM algorithm. The average metallicities for \citet{cald11} and \citet{kang12} are nearly consistent, while both the two groups from \citet{kang12} are metal-richer than the two groups from \citet{cald11}. Metallicities from \citet{kang12} show strong bimodal distribution at the 99.8\% confidence level, while the hypothesis of a unimodal distribution is rejected only at $\sim$64\% confidence level for metallicities from \citet{cald11}. Although many previous studies \citep{ab93,bh00,per02,fan08} have reported that the metallicity distribution in M31 is bimodal, \citet{cald11} suggested that the metallicity distribution in M31 is not generally bimodal, in strong distinction with the bimodal Galactic globular distribution. \begin{figure}[!htb] \figurenum{12} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig12.ps}}} \caption{Metallicity as a function of color. The open circles represent the confirmed GCs, while the red crosses represent the rest clusters, including candidate GCs, controversial objects, and extended clusters.} \label{fig:fig12} \end{figure} Figure 13 shows the relation between intrinsic colors and galactocentric distance $R_{\rm gc}$ for the sample clusters. Symbols are as in Figure 12. It can be seen that the colors of candidate GCs are on average redder than those of the confirmed GCs. We derived the mean color values in different projected galactocentric distance for the confirmed GCs. For clusters with galactocentric distance $R_{\rm gc}<$ 30 kpc and $R_{\rm gc}>$ 30 kpc, the mean values were measured with star clusters located within an annulus at every 3 kpc and 10 kpc radius from the center of M31, respectively, all of which were plotted with squares in Figure 13. \citet{crampton85} found that no radial $(B-V)_0$ color gradient exists for M31 GCs, however, \citet{sharov88} reported that the $(V-K)_0$ color shows a weak correlation with the galactocentric distance. No clear trend is present between the NIR colors and $R_{\rm gc}$ for the confirmed GCs. It seems that clusters with $R_{\rm gc}$ around 20 kpc are on average redder than inner clusters in the $(V-K_{\rm s})_0$ color index, however, this may be caused by the crowding blue clusters around the ``10 kpc ring'' \citep{Gordon06}, which pull the color index down. \begin{figure}[!htb] \figurenum{13} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig13.ps}}} \caption{Intrinsic color as a function of galactocentric distance $R_{\rm gc}$. Symbols are as in Fig. 12. The squares represent the mean color values in different projected galactocentric distance for the confirmed GCs.} \label{fig:fig13} \end{figure} The color-magnitude diagram (CMD) provides a qualitative model-independent global indication of cluster formation history \citep{ma12a, ma13}. Figure 14 displays the CMD of $M_{K_{\rm s}}$ vs. $(V-K_{\rm s})_0$ for the sample star clusters with magnitude uncertainty $< 0.5$ mag. Symbols are as in Figure 12. The absolute magnitudes were derived with the distance modulus of $(m-M)_0=24.47$ with a distance of $\sim$784 kpc. Several models were added to the CMD to obtain a more detailed history of cluster formation. Four fading lines from the simple stellar population (SSP) synthesis model \citet[][hereafter BC03]{bru03} for a metallicity of $Z=0.004$, $Y=0.24$, assuming a \citet{salp55} stellar IMF, and using the Padova-1994 isochrones, are plotted on the CMD of M31 star clusters for four different total initial masses: $10^6$, $10^5$, $10^4$, and $10^3$ $M_{\odot}$ (from up to bottom). It seems that many GC candidates in M31 are located out of the evolutionary tracks. The majority of M31 GCs fall between these four fading lines, consistent with previous results \citep{wang10}. \begin{figure}[!htb] \figurenum{14} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig14.ps}}} \caption{CMD of the sample clusters. Symbols are as in Fig. 12. Fading lines are indicated for clusters with total initial masses of $10^6$, $10^5$, $10^4$, $10^3$ $M_{\odot}$ (from up to bottom), assuming a Salpeter IMF.} \label{fig:fig14} \end{figure} Figure 15 shows the $(V-K_{\rm s})_0$ versus $(J-K_{\rm s})_0$ color-color diagram for M31 star clusters with magnitude uncertainty $< 0.5$ mag. Symbols are as in Figure 12. The theoretical evolutionary paths from the SSP model BC03 for $Z=0.004$, $Y=0.24$ (black line) and $Z=0.02$, $Y=0.28$ (green line) are displayed. The horizontal dashed line represents $V-K_{\rm s} = 3$, assuming the reddening value $E(B-V)=0.13$. As \citet{gall04} reported that, most of the background galaxies have $V-K_{\rm s} \geq3 $, providing a powerful tool to discriminate between M31 clusters and background galaxies. It can be seen that a large number of GC candidates are located above the dashed line, indicating that some background galaxies were mistaken for GC candidates. However, we should notice that the dashed line is just approximate to the criteria $V-K_{\rm s} = 3$, since one source with $V-K_{\rm s} > 3$ may be located below the dashed line if the reddening value is larger than 0.13 mag. The distribution of the clusters in the color-color diagram are much disperse, and many GC candidates are located out of the theoretical evolutionary paths, indicating that many GC candidates may not be true GCs. \begin{figure}[!htb] \figurenum{15} \center \resizebox{\hsize}{!}{\rotatebox{0} {\includegraphics{fig15.ps}}} \caption{Color-color diagram of the sample clusters. Symbols are as in Fig. 12. Theoretical evolutionary paths for Z = 0.004, Y = 0.24 (black line) and $Z$ = 0.02, $Y$ = 0.28 (green line) are drawn. The horizontal dashed line represents $V-K_{\rm s} = 3$, assuming the reddening value $E(B-V)=0.13$.} \label{fig:fig15} \end{figure} | 14 | 4 | 1404.4444 |
1404 | 1404.3052_arXiv.txt | Simultaneous radio and extreme ultraviolet (EUV)/white-light imaging data are examined for a solar type II radio burst occurring on 2010 March 18 to deduce its source location. Using a bow-shock model, we reconstruct the 3-dimensional EUV wave front (presumably the type-II emitting shock) based on the imaging data of the two STEREO spacecraft. It is then combined with the Nan\c{c}ay radio imaging data to infer the 3-dimensional position of the type II source. It is found that the type II source coincides with the interface between the CME EUV wave front and a nearby coronal ray structure, providing evidence that the type II emission is physically related to the CME-ray interaction. This result, consistent with those of previous studies, is based on simultaneous radio and EUV imaging data for the first time. | It has long been suspected that the interaction of a coronal mass ejection (CME) with nearby dense coronal structures like streamers or rays is important to solar type II radio bursts \citep[e.g.,][]{Wild1963, Classen2002, Reiner2003, Mancuso2004, Cho2005}. These coronal structures are featured by higher plasma density and lower bulk flow speed (in the magnetically-closed part of streamers) than that of the surrounding solar wind plasmas \citep[e.g.,][]{Habbal1997, Strachan2002}. Thus when a CME interacts with a nearby streamer or coronal ray, the associated disturbance may propagate into a region with much lower characteristic speed. This favors the formation or enhancement of coronal shock, as well as the consequent electron acceleration and type II excitation. Previous studies along this line of reasoning were largely based on a combined analysis of coronagraph images and radio spectrograph data. For example, \citet{Reiner2003} suggested that the type IIs of their study originated from dense region of the corona, likely from streamers. \citet{Cho2007} determined the shock heights by combining the type II spectral data from Green Bank Solar Radio Burst Spectrometer (GBSRBS; \citet{White2006}) and the polarization Brightness (pB) data from the Mauna Loa Solar Observatory (MLSO) MK4 coronameter \citep{Elmore2003}, and compared them to the MK4 CME-front heights. They concluded that the type II burst was generated at the interface of the CME flank and the streamer. In a follow-up study, \citet{Cho2008} expanded the study to 19 events and found that nearly half of them were probably associated with CME-streamer interaction. In a series of studies, \citet{Feng2012, Feng2013} and \citet{Kong2012} proposed a novel method to diagnose the type II sources by relating specific morphological features (e.g., bumps or breaks) of the dynamic spectra to imaging features (e.g., CME-driven shock propagating across a dense streamer). These studies confirmed that the CME-shock interaction with dense coronal structures like streamers is important to type II bursts. However, none of the studies was based on radioheliograph imaging data which provide the most direct evidence on the radio source location. This is due to the scarcity of simultaneous imaging in radio and white light/EUV wavebands. Nevertheless, existing studies have demonstrated the importance of combining these imaging observation as well as the spectral data in revealing the origin of metric type II bursts and their relationship with CME-driven shocks \citep[e.g.,][]{Bastian2001, Maia2000, Dauphin2006, Bain2012, Zimovets2012, Carley2013}. Along this line of observational endeavors, in this study we show a type II event with imaging data from both the Nan\c{c}ay radioheliograph (NRH; \citet{Kerdraon1997}) and instruments onboard the twin Solar TErrestrial RElations Observatory (STEREO; \citet{Kaiser2008}) and the Solar and Heliospheric Observatory (SOHO; \citet{Domigno1995}) spacecraft. This allows us to pinpoint the type II source location that is found to correspond to a CME-coronal ray interface. | \label{sec5} In this paper we show evidence supporting that the type II solar radio burst occurring on 2010 March 18 was originated from the interaction between a CME and a nearby dense coronal ray structure. This is based on simultaneous radio and EUV imaging data recorded at different vantage points. The NRH data were used to constrain the projected type II source location in the Earth FOV, and the multi-vantage point EUV data were used to reconstruct the 3-d profile of the EUV front (presumably the type-II emitting shock). We find that the type II source lay closely to the CME-ray intersection region. This study demonstrates the importance of CME interaction with dense coronal structures (like rays or streamers) to metric type IIs. The suspect that CME-streamer interactions may be important to metric type IIs \citep[e.g.][]{Wild1963, Classen2002, Reiner2003, Mancuso2004, Cho2005} stems from following considerations. Firstly, the CME-related disturbance may frequently run into a nearby density structure along the CME flank due to its lateral expansion. This interaction region likely corresponds to a quasi-perpendicular shock geometry, if a shock exists. Shocks of such geometry have been predicted to account for efficient electron acceleration \citep{Holman1983, Wu1984}. Secondly, it is expected that when a CME disturbance propagates into a dense coronal structure with low Alfv\'enic speed, the disturbance may steepen into a shock, or a preexisting shock may get enhanced in strength, also favoring electron acceleration. Thirdly, it is suggested that the closed field topology within the streamer, when swept by an outward-propagating shock, may give rise to a collapsing magnetic trap configuration \citep{Baker1987, Zlober1993, Somov1997, Chen2013}, within which electrons are trapped by the upstream field lines, and may return to the shock and be accelerated repeatedly. This also favors the generation of energetic electrons. While we do not consider the exact electron acceleration process in this word, our study is consistent with the above scenarios. | 14 | 4 | 1404.3052 |
1404 | 1404.3608_arXiv.txt | The recent high-statistics high-energy atmospheric neutrino data collected by IceCube open a new window to probe new physics scenarios that are suppressed in lower energy neutrino experiments. In this paper we analyze the IceCube atmospheric neutrino data to constrain the Violation of Equivalence Principle (VEP) in the framework of three neutrinos with non-universal gravitational couplings. In this scenario the effect of VEP on neutrino oscillation probabilities can be parametrized by two parameters $\Delta \gamma_{21}\equiv \gamma_2-\gamma_1$ and $\Delta\gamma_{31}\equiv \gamma_3-\gamma_1$, where $\gamma_i$'s denote the coupling of neutrino mass eigenstates to gravitational field. By analyzing the latest muon-tracks data sets of IceCube-40 and IceCube-79, besides providing the 2D allowed regions in $(\phi\Delta\gamma_{21},\phi\Delta\gamma_{31})$ plane, we obtain the upper limits $|\phi\Delta\gamma_{21}| < 9.1\times 10^{-27}$ (at 90\% C.L.) which improves the previous limit by $\sim4$ orders of magnitude and $|\phi\Delta\gamma_{31}| \lesssim 6\times 10^{-27}$ (at 90\% C.L.) which improves the current limit by $\sim1$ order of magnitude. Also we discuss in detail and analytically the effect of VEP on neutrino oscillation probabilities. | The Equivalence Principle is the cornerstone of classical gravitational theories, from Newtonian gravitation to General Relativity. The Weak Equivalence Principle (WEP) states that the geodesic paths followed by free falling bodies are the same, regardless of their energy content. In the other words, the motion of a falling body is determined only by the surrounding geometry and not by the body's own properties~\cite{Gravitation}. In the weak field limit, this principle leads to an universal acceleration of the falling bodies, a fact that is rooted in two principles of the Newtonian gravitation: the equivalence of inertial and gravitational masses and universality of the Newton's gravitational constant $G_N$. Since the proposal of WEP, this hypothesis has been extensively tested by a large diversity of experiments, including torsion-balance experiments~\cite{Eot-wash}, motion of solar system bodies~\cite{Overduin:2013soa}, spectroscopy of atomic levels~\cite{Hohensee:2013cya} and pulsars~\cite{Damour:1991rq,Horvat:1998st,Barkovich:2001rp}; which always lead to strong limits on possible deviations. However, recent developments in theoretical physics are systematically indicating that many modern attempts to obtain a quantum version of the gravitational theory lead to the prediction that the equivalence principle will be violated in some scale (see for example~\cite{Damour:2001fn,Damour:2002mi,ArmendarizPicon:2011ys,Damour:2010rm,Carroll:2008ub,Olmo:2006zu,Adunas:2000zm}). In this sense, improving the current limits on the VEP provides a diagnostic tool in probing very high energy theories of quantum gravity, which are almost inaccessible to conventional experiments. One of the methods to probe VEP is through the neutrino oscillation phenomena. The effect of VEP on neutrino oscillation was first studied by Gasperini~\cite{Gasperini:1988zf} and later developed by others in~\cite{Gasperini:1989rt,Halprin:1991gs}. The original model was intended to solve the solar neutrino problem~\cite{Pantaleone:1992ha,Butler:1993wi,Bahcall:1994zw,Halprin:1995vg,Mureika:1996ud,Mansour:1998nb,Gago:1999hi,Casini:1999kt,Majumdar:2000sd}, which is now in excellent agreement with the framework of massive neutrinos with the MSW effect~\cite{Mikheev:1986wj,Mikheev:1986gs}. However, despite its failure to explain the solar neutrino problem, VEP can contribute to flavor oscillation as a subdominant effect and so can be probed by solar neutrinos~\cite{Minakata:1994kt,Valdiviesso:2011zz}, atmospheric~\cite{Foot:1997kk,Foot:1998pv,Foot:1998vr,Fogli:1999fs,GonzalezGarcia:2004wg,GonzalezGarcia:2005xw,MACRO-ref,Morgan:2007dj,Abbasi:2009nfa}, supernova~\cite{Pakvasa:1988gd,Guzzo:2001vn}, cosmic~\cite{Minakata:1996nd} and accelerator~\cite{Iida:1992vh,Mann:1995nw} neutrinos. Essentially, the sensitivity of neutrino oscillation to VEP originates from the fact that the flavor states of neutrinos are coherent superposition of mass eigenstates and so act as interferometers which are sensitive to differences in the coupling of mass states to gravitational field. The bottom line is that VEP effectively changes the mass-squared differences by adding a term proportional to the square of neutrino energy ($\propto E_\nu^2$). Thus, by the increase of neutrino energy the VEP effects become stronger and so the potential to discover/constrain VEP increases. Among the known perpetual sources of neutrinos, atmospheric neutrino energies extends up to very high energy and so provides a unique opportunity to probe VEP. The construction of huge (km$^3$ scale) neutrino telescopes, with the completed IceCube detector at the South Pole as an example, fulfills the detection of these high energy atmospheric neutrinos. Currently two sets of high energy atmospheric neutrino data are available from IceCube experiment: the ``IC-40" data set in the energy range (100 GeV - 400 TeV)~\cite{Abbasi:2010ie} and ``IC-79" data set in the range (20 GeV - 10 TeV)~\cite{Aartsen:2013jza}, with the total number of events: $\sim18,000$ and $\sim40,000$ respectively. In this paper we utilize these data in the search of VEP in the most general phenomenological model accommodating it. By analyzing these data we obtain the most stringent upper limit of VEP parameters, some of them are $\sim4$ orders of magnitude stronger than the current limits. This paper is organized in the following way: in section~\ref{model} we review the phenomenology of the oscillation of massive neutrinos in the presence of VEP and current upper limits on VEP parameters. In section~\ref{sec:vep-energy} we study in detail the effect of VEP on neutrino oscillation. Also, we show the numerical calculation of oscillation probabilities and their interpretation in terms of analytical approximations. Our analysis of the atmospheric neutrino data of IceCube is presented in section~\ref{analysis}. Conclusion is provided in section~\ref{sec:con}. | \label{sec:con} One of the essential pillars in the theory of gravitation, both classical and relativistic, is the equivalence principle which has been tested in a variety of experiments. Violation of equivalence principle has far-reaching consequences in the neutrino sector, basically introducing novel oscillation pattern which can be measured at neutrino oscillation experiments. Thus, neutrino phenomenology provides a unique tool to probe the possible violation of equivalence principle. The strength of VEP effect on neutrino oscillation depends on the neutrino energy: the VEP effectively introduce mass-squared differences proportional to $E_\nu^2$ and so the oscillation phase will be proportional to $E_\nu$. Thus, clearly, the recent collected data of high energy ($\gtrsim100$~GeV) atmospheric neutrinos by IceCube experiment can discover/constrain VEP unprecedentedly. In this paper we studied the effect of VEP on the oscillation of high energy atmospheric neutrinos. In the high energy range the conventional standard oscillation induced by $\Delta m_{ij}^2$ is absent and the survival probability of each neutrino flavor is $\sim1$. However, VEP can drastically change this pattern: the effective energy-dependent mass-squared differences induced by VEP can lead to resonance flavor conversions and also oscillatory behavior in high energy with new maxima and minima in flavor oscillation probabilities. For the phenomenological model of VEP we considered in this paper, with the two VEP parameters $\phi\Delta \gamma_{21}$ and $\phi\Delta \gamma_{31}$, we studied in detail the oscillation pattern and provided the analytical descriptions of oscillation probabilities. We justified the numerical calculation of oscillation probabilities (especially $P(\nu_\mu\to\nu_\mu)$ which plays the main role in IceCube analysis) with the obtained analytical expressions and showed that the analytical approximation explains the oscillation pattern with impressive accuracy. Furthermore, we confronted the expected zenith distribution of muon-track events in the presence of VEP with the collected data by IceCube experiments with two different configurations, namely IC-40 and IC-79 data sets. To analyze these data we performed a simple $\chi^2$ analysis taking into account the statistical and systematic errors. The oscillation probabilities have been calculated numerically by scanning the parameter space of VEP parameters in the full three flavors framework. From these analyses we obtained the following bounds on the VEP parameters at 90\% C.L.: $-9.2\times10^{-27} < \phi\Delta\gamma_{21} < 9.1\times10^{-27}$ and $-6.3\times10^{-27} < \phi\Delta\gamma_{31} < 5.6\times10^{-27}$. The obtained limit on $\phi\Delta\gamma_{21}$ is $\sim4$ orders of magnitude stronger than the current limit; also we improved the existing bound on $\phi\Delta\gamma_{31}$ by $\sim1$ order of magnitude. Finally we investigated the future sensitivity of IceCube to VEP parameters. We have presented the sensitivity region in VEP parameter space assuming three times of IC-79 data set; which improves mildly the obtained limits. Also, we have studied the effect of VEP on cascade events in IceCube, motivated by the fact that VEP induces $\nu_\mu\to\nu_e$ conversion that can distort the energy distribution of cascade events. However, due to lower statistics and higher uncertainties for cascade detection, the sensitivity of IceCube to VEP parameters in cascade channel is less than the sensitivity in muon-track channel. At the end we would like to emphasize that the limits obtained in this paper can be translated to limits on the parameters of theories (either effective theories or extensions of Standard Model) which predict/accommodate VEP to some level. As an example in this line we can mention the Standard Model Extension (SME) theories which consist of extending the Standard Model action by including all the possible terms that violate the Lorentz invariance~\cite{Colladay:1996iz,Colladay:1998fq,Kostelecky:2010ze}. One of the consequences of SME is the violation of equivalence principle such that test of VEP provide a tool for searches of Lorentz symmetry violation. Further speculations regarding these connections and possibilities to probe fundamental theories by VEP tests in neutrino sector~\cite{Kostelecky:2003cr,Kostelecky:2003xn,Kostelecky:2004hg,Diaz:2009qk,Katori:2006mz} will be pursued in a later work. | 14 | 4 | 1404.3608 |
1404 | 1404.1920_arXiv.txt | We combine a new, comprehensive database for globular cluster populations in all types of galaxies with a new calibration of galaxy halo masses based entirely on weak lensing. Correlating these two sets of data, we find that the mass ratio $\eta \equiv M\sbr{GCS}/M\sbr{h}$ (total mass in globular clusters, divided by halo mass) is essentially constant at $\langle \eta \rangle \sim 4 \times 10^{-5}$, strongly confirming earlier suggestions in the literature. Globular clusters are the only known stellar population that formed in essentially direct proportion to host galaxy halo mass. The intrinsic scatter in $\eta$ appears to be at most 0.2 dex; we argue that some of this scatter is due to differing degrees of tidal stripping of the globular cluster systems between central and satellite galaxies. We suggest that this correlation can be understood if most globular clusters form at very early stages in galaxy evolution, largely avoiding the feedback processes that inhibited the bulk of field-star formation in their host galaxies. The actual mean value of $\eta$ also suggests that about $1/4$ of the \emph{initial} gas mass present in protogalaxies collected into GMCs large enough to form massive, dense star clusters. Finally, our calibration of $\langle \eta \rangle$ indicates that the halo masses of the Milky Way and M31 are $(1.2\pm0.5)\ten{12} M_{\sun}$ and $(3.9\pm1.8)\ten{12} M_{\sun}$ respectively. | In the standard paradigm of galaxy formation, dark matter halos form and merge through gravitational instability, and within these halos gas cools and stars form. An essential ingredient of galaxy formation models is ``feedback'' which injects energy or momentum into the gas, either heating it or driving it outward. Commonly cited feeback mechanisms include supernovae and stellar winds in small galaxies, and active galactic nuclei (AGN) and infall gas heating in large galaxies \citep[see e.g.\ the review by][]{SilMam12}. The bulk of the star formation and AGN activity, and hence presumably the bulk of the feedback, occurs at redshifts $z \sim 1$ -- $3$, corresponding to lookback times 8 -- 11.5 Gyr \citep{BehWecCon13}. Empirically star formation is most efficient for galaxies with halo masses in the intermediate range $10^{12}-10^{13} \msun$, for which 20-25\% of the baryons are converted into stars \citep{MarHud02, LeaTinBun12, BehWecCon13, VelvanHoe14, HudGilCou13}. Globular clusters (GCs) have been found in virtually every galaxy from dwarfs to brightest cluster galaxies (BCGs), excluding only the very tiniest dwarfs. They are commonly regarded as relics of the earliest star-forming stages in their host galaxies, a view long confirmed by direct measurements of their ages \citep[see, e.g.][]{VanBroLea13}. While in some galaxies affected by \emph{late} gas-rich mergers, younger GCs are present \citep[][and later papers]{AshZep92}, the formation times for \emph{most} GCs belong to the redshift range $z \sim 2 - 8$ and thus precede the bulk of star formation in the Universe. The first GCs likely started forming within the pregalactic gas-rich and nearly pristine dwarfs, out of dense, particularly massive GMCs (giant molecular clouds) \citep[e.g.][among others]{Har94, har10}. Growing evidence also suggests that the GCs were likely to have formed a little before the bulk of the metal-poor field-star population \citep[e.g.][]{Bla99,har02,pen08,Spi10}. This ``head start'' on formation, plus their intrinsically dense and massive structure, can be expected to have made them less easily influenced by feedback processes even as protoclusters, than were the field stars within their host galaxies. Moreover, the individual GCs in all kinds of galaxies exhibit almost identical distributions by mass, heavy-element abundance, and King-model-type structures \citep[see][for a review]{har10}. Globular cluster \emph{systems} (GCSs; the ensembles of all the GCs in a given galaxy) thus provide observers with a direct view into a remarkable common thread of galaxy formation history in its initial stages. Because most GCs started forming before most of the stars in the Universe, their formation may relate in a simpler way to the dark matter distribution than does the bulk of the stellar mass in galaxies. \cite{BlaTonMet97} first suggested that the total numbers of GCs in BCGs were directly proportional to the total, dark-matter dominated mass of the galaxy cluster. This suggestion was extended to host galaxies of all types especially by \cite{spi09}, \cite{geo10}, and \cite{hha13} among others. Those studies used halo mass estimates based primarily on models and relatively small galaxy datasets, and spliced together a variety of methods to cover the entire range of galaxy masses. Suggestions based on galaxy formation simulations that the GC population should correlate closely with the dark matter potential have also been made by \citet{KraGne2005} and \citet{MooDieMad2006}, though for somewhat different reasons. In this paper, we take advantage of a large new comprehensive GCS database to derive $M\sbr{GCS}$, the total mass of all globular clusters in a given galaxy. This dataset spans five orders of magnitude in galaxy stellar mass, and includes every galaxy type from dwarfs to supergiants. We combine these with a new observationally-based prescription for calculating galaxy halo mass, $M\sbr{h}$, based entirely on weak lensing. In agreement with the papers cited above, we find that the simple ratio $\eta \equiv M\sbr{GCS}/M\sbr{h}$ \citep{geo10} is virtually constant over the entire range, very much unlike the strongly nonlinear behavior of total \emph{stellar} mass $M_{\star}$ with $M\sbr{h}$. An outline of this paper is as follows. After providing the background for the data in section 2, we discuss the distribution of $\eta$ in section 3. In section 4 we discuss possible factors contributing to the observed scatter, and the implications of the relation for galaxy and GMC formation. | The residual scatter in $\eta$ is 0.3 dex when galaxies with the best-estimated GC counts (uncertainties $\sigma(\log M\sbr{GCS}) < 0.1$ dex) are used. Because the halo masses $M\sbr{h}$ are not direct measurements but are inferred from the stellar mass and the mean SHMR, the scatter in the SHMR cannot be calculated directly from weak lensing, but by other methods, $\sigma(M\sbr{h})$ has been estimated at 0.15 to 0.2 dex \citep{BehConWec10}. Subtracting 0.2 dex in quadrature then suggests that the intrinsic scatter in the $M\sbr{GCS}$--$M\sbr{h}$ relation is a remarkably low $\sim 0.2$ dex. Among several potential sources for the scatter about the mean $M\sbr{GCS}$--$M\sbr{h}$ relation, we discuss two such sources: tidal stripping of the GC system as a whole, and the red and blue GC subpopulations. \subsection{Environmental Dependence and Tidal Stripping of the Globular Cluster System} At large stellar or halo masses, the data are suggestive of an environmental dependence in the sense that non-BCGs have slightly lower $\eta$ than BCGs. One obvious interpretation is that the GC system has been partially stripped in galaxies which are ``satellites'', as opposed to ``centrals'' in their host halo \citep{spi09}. The Coma cluster may again provide a good illustration: the two supergiants NGC 4874 and 4889 dominate the Coma center and have comparable stellar mass. But from dynamical and X-ray studies \citep{ColDun96, ArnAghGas01}, the presence of a cD halo, and the spatial distribution of intra-cluster GCs \citep{PenFerGou11}, it is clear that NGC 4874 is currently the BCG of the Coma cluster. At some earlier time, NGC 4889 was presumably the dominant galaxy of its own poor cluster, before it merged with a comparable-sized cluster centered on NGC 4874 to form what is now the Coma cluster. If NGC 4889 is orbiting the potential well dominated by NGC 4874, then we expect that its own halo of dark matter and GCs has at least partially been stripped off by tidal effects, and has joined the extended Coma cluster halo. Of course the degree of stripping depends on how tightly bound a given population is. The dark matter, which is a very extended population, should be most easily stripped \citep[$\sim 65$\% retained,][]{GilHudErb13}, while the stellar light is tightly bound to the infalling galaxy and is most difficult to strip. But what is stripped would become the intracluster light. The GCS is, on average, more compact than the dark matter, but less so than the field-star light and so would be more easily stripped than the latter \cite[see][for stripping in dwarf galaxies]{SmiSanFel13}. As discussed in Section 3, the halo mass for satellites that appears in the denominator of $\eta$ is the halo mass {\emph before} stripping, whereas the numerator is the GCS after stripping. Consequently, the $\eta-$values for the non-BCG satellites would be reduced \citep[see also][]{spi09}. \subsection{Red vs. Blue GCs} GCSs commonly have \emph{bimodal} distributions in GC color or metallicity \citep[e.g.][among a host of others] {lar01, pen06, har09a, mie10}; `red' GCs have mean [Fe/H] $\simeq -0.5$, while `blue' GCs have mean [Fe/H] $\simeq -1.5$. These two subpopulations are thought to have formed at different stages of hierarchical merging, with additional metal-poor ones accreted later from satellite dwarfs \citep[][among others]{Har94, bur01, har10, ton13}. Direct age measurements indicate significant overlap in their age distributions, with the more metal-rich ones only $1-2$ Gyr younger than the metal-poor ones on average \citep[e.g.][] {LeaVanMen2013,han13}. The metal-richer clusters are almost absent in dwarfs but become progressively more prominent with galaxy luminosity, growing to roughly half the total population in giants \citep{pen06, pen08}. However, the exact blue/red proportions differ from galaxy to galaxy even at the same luminosity, and such differences may contribute to the scatter in Figure 1. A direct test would be to plot $\eta$ separately for the blue and red subpopulations and see how the residual scatter changes \citep{spi09}. Unfortunately, the numbers for $N\sbr{GCS}$ (blue, red) are not available yet for most galaxies in the catalogue of \cite{hha13}, so thorough comparisons must await future work. \subsection{Implications for GC and galaxy formation} The empirical fact that $M\sbr{GCS}$ is a nearly constant fraction of $M\sbr{h}$ must have its origin in early star-forming conditions. We propose that $\eta \simeq$ constant can result if three conditions are met: \begin{enumerate} \item the \emph{initial gas mass} present in a pregalactic potential well is proportional to halo mass; \item the \emph{globular cluster formation rate} is proportional to the available gas mass; and \item globular clusters form early and before feedback effects such as stellar winds and supernovae (for dwarf galaxies), and AGN activity and infall heating (for giant galaxies) begin inhibiting star formation. \end{enumerate} We suggest also that the approximate value for the ``absolute'' efficiency ratio ${\eta}$ can be understood as the product of four separate mass ratios: \begin{align} \eta & \sim \Bigl(\frac{M\sbr{bary}}{{M\sbr{h}}}\Bigr)\times \Bigl(\frac{M\sbr{GMC}}{M\sbr{bary}}\Bigr)\times \Bigl(\frac{M\sbr{PGC}}{M\sbr{GMC}} \Bigr)\times\Bigl(\frac{M\sbr{GC}}{M\sbr{PGC}}\Bigr) \nonumber \\ ~ & \sim 0.15 \times \Bigl(\frac{M\sbr{GMC}}{M\sbr{bary}}\Bigr) \times 0.01 \times 0.1 ~ \sim 4\times 10^{-5} \end{align} Here, $M\sbr{bary}$ is the total baryonic mass in the pregalactic halo (by hypothesis, the total initial gas mass); $M\sbr{GMC}$ is the total gas mass that collects into GMCs; $M\sbr{PGC}$ is the gas mass within a protoglobular cluster as it begins star formation; and finally $M\sbr{GC}$ is the present-day mass of the globular cluster. The first term in equation (1) is the typical universal baryonic-to-dark mass ratio \citep{Planck13-16} and is $0.15$. For the third term, observational evidence \citep{Har94, LadLad2003} suggests that of order 1\% of the turbulent gas in a GMC will clump into the especially dense protoclusters that will produce massive star clusters capable of surviving over the long term. Lastly, for the fourth term, a present-day GC is typically about 10\% as massive as its initial (gaseous) protocluster because (a) the star formation efficiency within the PGC should be $\sim 0.3 - 0.5$, and (b) over 12 Gy of dynamical evolution in its host galaxy, the GC will lose 1/3 or more of its initial \emph{stellar} mass due to early rapid evolution of massive stars and later, slower tidal stripping and evaporation \citep[e.g.][]{LadLad2003,Kru2013,WebHarSil2013}. The second term $(M\sbr{GMC}/M\sbr{bary})$ is the most difficult to estimate: it represents the fraction of all gas \emph{at the epoch of GC formation} that succeeds in cooling and collapsing into GMCs large enough to form globular clusters. Instead we can use our direct calibration of $\langle \eta \rangle$ and invert eq.~(1) to derive \begin{equation} \Bigl(\frac{M\sbr{GMC}}{M\sbr{bary}}\Bigr) \sim 0.25 \end{equation} This suggests that GMC formation in the protogalaxies was considerably more efficient than in present-day $L_{*}$ spiral galaxies for which $M\sbr{H_{2}}/M\sbr{bary} \sim 0.01$ \citep{BosCorBoq2014}. \subsection{Future Prospects} Several directions for future work will lead to a better understanding of the $M\sbr{GCS}-M\sbr{h}$ correlation. As next stages, we are searching the \cite{hha13} database for galaxies with both wide-field and multicolor data which will define $\eta$ separately for the blue and red GCs. Beyond that it will be necessary to make new observations, preferably for a wider range in galaxy size, type and environment. The correlation of $\eta$ with environment discussed in Section 4.1 suggests that tidal stripping plays an important role for satellite galaxies in the group/cluster environment. If we assume a universal $\eta$, then the observed $\eta$ could be used to measure the tidal stripping of the GCSs of individual satellite galaxies as a function of their host (group or cluster) halo mass and of their orbital properties \citep{OmaHudBeh13}. | 14 | 4 | 1404.1920 |
1404 | 1404.2112_arXiv.txt | We present a high-precision, differential elemental abundance analysis of the HAT-P-1 stellar binary based on high-resolution, high signal-to-noise ratio Keck/HIRES spectra. The secondary star in this double system is known to host a transiting giant planet while no planets have yet been detected around the primary star. % The derived metallicities ([Fe/H]) of the primary and secondary stars are identical within the errors: $0.146 \pm 0.014$ dex ($\sigma$ = 0.033 dex) and $0.155 \pm 0.007$ dex ($\sigma$ = 0.023 dex), respectively. Extremely precise differential abundance ratios of 23 elements have been measured (mean error of $\sigma$([X/Fe]) = 0.013 dex) and are found to be indistinguishable between the two stars: $\Delta$[X/Fe] (secondary - primary) = +0.001 $\pm$ 0.006 dex ($\sigma$ = 0.008 dex). The striking similarity in the chemical composition of the two stellar components in HAT-P-1 is contrary to the possible 0.04 dex level difference seen in 16 Cyg A+B, which also hosts a giant planet, at least 3 times more massive than the one around HAT-P-1 secondary star. We conclude that the presence of giant planets does not necessarily imply differences in the chemical compositions of the host stars. The elemental abundances of each star in HAT-P-1 relative to the Sun show an identical, positive correlation with the condensation temperature of the elements; their abundance patterns are thus very similar to those observed in the majority of solar twins. In view of the \citet{m09}'s interpretation of the peculiar solar abundance pattern, we conclude that HAT-P-1 experienced less efficient formation of terrestrial planets than the Sun. This is in line with the expectation that the presence of close-in giant planets preventing the formation or survival of terrestrial planets. | The components of binary systems are usually assumed to share the same origins and to have identical chemical compositions. Several studies, however, find that the elemental abundance differences in stellar binaries are in fact not uncommon \citep{g01,lg01,d04,d06,r11}. The sources of these abundance deviations remain unknown but one possible explanation is that any elemental abundance differences could relate to the processes of planet formation. % It is well established that the dwarfs and sub-giants with higher metallicity have higher probability to form giant planets \citep{g97,fv05,us07,j10} yet the situation for the giants is still uncertain \citep{p07,h07,t08,m13}. Meanwhile whether planet formation can affect the chemical compositions of the host stars is less clear. \citet{m09} found small but statistically significant anomalies in the solar chemical composition compared to most solar twins and argued that these were due to the formation of terrestrial planets in the solar system that preferentially locked-up refractory elements (i.e. those easily condensing); the deficiency of refractory elements in the solar photosphere would disappear if some four Earth-masses of terrestrial planet material would be added into the present solar convective envelope \citep{c10}. \citet{m09} also found that the stars found {\em not} to have a close-in giant planet are more likely to resemble the Sun chemically, which suggests that the presence of close-in giant planets might prevent the formation of terrestrial planets. One could also speculate that in these systems, the smaller planets were accreted by the host stars during the migration process of Jupiter-like planets, thus removing the initial stellar abundance signature imprinted by the process of planet formation. \citet{l11} provide observational support for this idea based on Kepler data. \citet{r11} demonstrated metallicity differences in the binary system 16 Cyg A+B to be $0.04 \pm 0.01$ dex (16 Cyg A is more metal-rich than 16 Cyg B) and related it to planet formation; 16 Cyg B is known to host a giant planet with a minimum mass of 1.68 Jupiter masses \citep{c97}. On the other hand, \citet{s11} found no such abundance differences in 16 Cyg A+B. The reasons for these contrary results remain unknown but given the possible connection between planet formation and stellar host composition, there is an urgent need for additional binary systems hosting planets to be exposed to a high precision abundance analysis. Here we present such a study for the HAT-P-1 stellar binary, in which a close-in giant planet orbits around the secondary star \citep{b07}; no planet detection around the primary star has been reported. | \begin{figure} \includegraphics[width=\columnwidth]{Fig2.eps} \caption{Differential elemental abundances of HAT-P-1 stellar binary relative to our solar abundances and to each other as a function of dust condensation temperature; filled circles and blue triangles represent [X/Fe] without and with GCE corrections, respectively. Black dashed lines and blue dot-dashed lines show the fitting slopes of our results without and with GCE corrections, respectively. Green solid lines show the mean trend of solar twin stars according to \citet{m09}.} \label{f:tcondensation} \end{figure} The focus of this discussion is to examine the possible connection between planet formation and stellar host composition in the HAT-P-1 binary. Given that the secondary star in the HAT-P-1 stellar binary is known to harbor a giant planet, our high precision chemical abundances may place new constraints on the planet formation process, at least in this system. As noted in the introduction, \citet{m09} and follow-up studies \citep{r09,r10} discovered that the Sun shows deficiency in refractory elements relative to volatiles when compare to the majority ($\sim 80-90$\%) of solar twins. The deficiencies correlate with the condensation temperature ($T_{\rm c}$) of the elements such that the abundances of refractory elements ($T_{\rm c} \geq$ 900 K) decrease (Sun - solar twins) with increasing $T_{\rm c}$. They argue that the special abundance pattern of the Sun is due to dust condensation and terrestrial planet formation in the proto-solar disk that for some reason proceeded more efficiently than for the majority of solar twins. They then argue that terrestrial planets over giant planets as the cause for the peculiar abundance signature due to the presence of a break at $T_{\rm c} \approx 1200$\,K (much higher than the expected temperatures in the proto-planetary disk where the solar system giant planets formed), the required amount of refractory material necessary to imprint the signature ($4$\,M$_\oplus$) and the higher frequency of stars sharing the solar abundance pattern that do {\em not} have a close-in giant planet. We note however that \citet{o14} propose that the abundance differences found by \citet{m09} are not the result of planet formation but are imprinted by dust-gas separation in the interstellar medium prior to star formation based on their finding that all of their 14 stars in the open cluster M67 resemble more the Sun than the solar twins of \citet{m09}. They conclude that the Sun formed in an unusually dense stellar environment like M67. The existence of a high temperature break in $T_{\rm c}$ and the apparent correlation with absence of close-in giant planets are not easily understood in that scenario however. Our high quality data allow us to make robust conclusions about the [X/Fe] -- $T_{\rm c}$ slopes of the HAT-P-1 stellar binary. Fig. \ref{f:tcondensation} shows the differential abundances of HAT-P-1 primary and secondary star relative to the Sun and relative to each other versus $T_{\rm c}$ \citep{l03}\footnote{[O/Fe] of both stars relative to the Sun would fall to the fitting trends while $\Delta$[O/Fe] (relative to each other) would be 0.03 dex larger if NLTE corrections are adopted from \citet{f09}.}. All the elemental abundances were used to derive the slopes. The slopes of linear fitting for both stars compared to the Sun are positive and identical within errors: $(1.15 \pm 0.10) \times 10^{-4}$ dex K$^{-1}$ and $(1.28 \pm 0.08) \times 10^{-4}$ dex K$^{-1}$ for primary and secondary star, respectively. These slopes are very similar to the trends of refractories of the average of solar twins relative to the Sun \citep{m09,r09}, which with their interpretation would imply that both binary components formed less terrestrial planets than the Sun. This is consistent with the expectation that the presence of close-in giant planets prevents the formation or survival of terrestrial planets \citep{il04}. The positive slopes can also arise from Galactic chemical evolution (GCE). Therefore we applied the GCE corrections on our [X/Fe] values based on the studies of \citet{gh13}. We adopted the \citet{gh13}'s data and fitting trends to derive the values of [X/Fe] at [Fe/H] $\sim$ 0.15 dex to correct our results. The final results with GCE corrections only show tiny differences of the general trends (see Fig. \ref{f:tcondensation}a,b) which indicate that these positive slopes can not be erased even after GCE corrections. When comparing the two HAT-P-1 components relative to each other, the slope of [X/Fe] -- $T_{\rm c}$ is non-existent (Fig. \ref{f:tcondensation}c): $(0.60 \pm 6.36) \times 10^{-6}$ dex K$^{-1}$. As stated before, the mean elemental abundance difference between the secondary and primary star is +0.001 $\pm$ 0.006 dex ($\sigma$ = 0.008 dex, secondary - primary). Clearly, the two stars have indistinguishable chemical compositions, which is interesting given the detection of a close-in giant planet with mass $\sim 0.53\,{\rm M_{Jupiter}}$ around the secondary star. We conclude that the formation process of giant planets does not necessarily affect the chemical pattern of the host star, which supports the conclusions of \citet{m09} and \citet{r09}. This is contrary to the difference of $0.04 \pm 0.01$\,dex seen in 16 Cyg A+B \citep{r11} but is consistent with the results from \citet{s11}. Assuming for the moment that the 16 Cyg abundance differences are real, one possible explanation could be the higher mass ($2.4\,{\rm M_{Jupiter}}$) of the 16 Cyg planet \citep{p13}: it is still possible that such a more massive planet imprints a chemical signature in the host star. We note, however, the stellar masses in HAT-P-1 are slightly higher ($1.16$ and $1.12$\,$M_{\sun}$, \citealp{b07}) than in 16 Cyg ($1.05$ and $1.00$\,$M_{\sun}$, \citealp{r11}), which makes the convection zone less massive and thus more prone to chemical imprints from planet formation. Albeit the smaller convection zone in HAT-P-1 would make it easier to imprint a planet signature, higher mass stars seem to have shorter disk lifetimes \citep{w11}, making thus more difficult to imprint any planet signature in HAT-P-1 than in 16 Cyg. Which of these effects that dominate would depend on the exact size of the convection zone at the time of the accretion and the amount of material heavier than Helium locked up in the giant planet. For the time being, we conclude that the formation of giant planets do not necessarily have to introduce chemical signatures in their host stars. Our detailed study of the HAT-P-1 double system underscores how high precision differential abundance measurements in binary stars with planets can provide important constraints on planet formation. Further efforts are needed to examine the physical characteristics and chemical abundances for additional stellar binaries with giant or terrestrial planets in order to understand the formation and evolution of planetary systems. | 14 | 4 | 1404.2112 |
1404 | 1404.7501_arXiv.txt | *{} \abstract{Eclipsing binary stars have long served as benchmark systems to measure fundamental stellar properties. In the past few decades, asteroseismology - the study of stellar pulsations - has emerged as a new powerful tool to study the structure and evolution of stars across the HR diagram. Pulsating stars in eclipsing binary systems are particularly valuable since fundamental properties (such as radii and masses) can determined using two independent techniques. Furthermore, independently measured properties from binary orbits can be used to improve asteroseismic modeling for pulsating stars in which mode identifications are not straightforward. This contribution provides a review of asteroseismic detections in eclipsing binary stars, with a focus on space-based missions such as CoRoT and \kep, and empirical tests of asteroseismic scaling relations for stochastic (``solar-like'') oscillations.} | Asteroseismology has undergone a revolution in the past few decades. Driven by multi-site ground-based observing campaigns and high-precision space-based photometry, the number of stars with detected pulsations has increased dramatically, and pulsation frequencies and amplitudes are measured with unprecedented precision. In particular the photometric data provided by the CoRoT and \kep\ space telescopes have allowed the application of asteroseismology to stars throughout the HR diagram \citep[e.g.,][]{gilliland10,michel12,chaplin13b}. Owing to the relatively large apertures of some space-based telescopes, however, the majority of stars with high quality asteroseismic detections are relatively faint, and hence lack independent observational constraints from classical methods such as astrometry or long-baseline interferometry. Combining independent observations with asteroseismology is crucial to advance progress in theoretical modeling of observed oscillation frequencies, and the validation of asteroseismic relations to derive fundamental stellar properties. Therefore, the full potential of asteroseismology can only be realized if such observations can be combined with model-independent constraints on properties such as temperatures, radii and masses. Eclipsing binaries have long served as benchmark systems for determine fundamental properties of stars from first principles. Similar to measuring oscillation frequencies, the observation of photometric eclipses and spectroscopic radial velocities can be performed for relatively faint systems, as long as spectral lines of the components can be successfully disentangled. Furthermore, eclipses and pulsations can be measured using the same data. Thus, asteroseismology of components in eclipsing binary systems promises to be a powerful method to improve our understanding of stellar structure and evolution. | Table \ref{tab:summary} summarizes the characteristics of confirmed eclipsing and/or heartbeat systems for which asteroseismic detections (either self-driven or tidally-induced) have been made using space-based observations. The list illustrates that the synergy of asteroseismology and eclipsing/eccentric binary stars using space-based observations is (unsurprisingly) still in its infancy: all of the discussed systems have been published within the last four years. Most detections were made by \kep, which provided the required continuous monitoring to detect eclipses, and high-precision photometry to detect oscillations despite the dilution by the binary component or small amplitudes in the pulsating star. As noted in Section 4.1, the list of confirmed red giants in eclipsing systems can be expected to increase significantly once sufficient radial velocities have been gathered for the \kep\ candidate systems \citep{gaulme13}. \begin{table*} \begin{small} \begin{center} \caption{List of eclipsing and/or heartbeat systems with components showing self-excited or tidally-induced pulsations detected from space-based observations. Systems are grouped into giants (top), intermediate and high-mass stars (middle) and compact stars (bottom). Columns list the approximate $V$-band magnitude, spectral types (Sp.Type), orbital period ($P$), eccentricity ($e$), and flags indicating a double-lined spectroscopic orbit (SB2), stochastic oscillations (stoch.), self-excited coherent pulsations, tidal pulsations, eclipses (Ecl.), and heartbeat effects (HB).} \begin{tabular}{l c c c c c c c c c c c} \hline ID & $V$ & Sp.Type & $P$(d) & $e$ & SB2 & stoch. & coherent & tidal & Ecl. & HB & Ref \\ \hline HD\,181068 & 8.0 & KIII+MV+MV& 45.5+0.91 & 0.0 & no & no & no & yes & yes & no & a \\ KIC\,5006817 & 10.9 & KIII+MV & 94.8 & 0.71 & no & yes & no & no & no & yes & b \\ KIC\,8410637 & 11.3 & KIII+FV & 408.3 & 0.69 & yes & yes & no & no & yes & no & c \\ \hline KIC\,10661783 & 9.5 & GIV+AV & 1.2 & 0.0 & yes & no & yes & no & yes & no & d \\ KIC\,4544587 & 10.8 & FV+FV & 2.2 & 0.29 & yes & no & yes & yes & yes & yes & e \\ HD\,174884 & 8.4 & BV+BV & 3.7 & 0.29 & yes & no & maybe & maybe & yes & yes & f \\ CID\,102918586 & 11.7 & FV+FV & 4.4 & 0.25 & yes & no & yes & yes & yes & yes & g \\ KIC\,11285625 & 10.1 & FV+FV & 10.8 & 0.0 & yes & no & yes & no & yes & no & h \\ KIC\,3858884 & 9.3 & FV+FV & 26.0 & 0.47 & yes & no & yes & maybe & yes & no & i \\ HD\,187091 & 8.4 & AV+AV & 41.8 & 0.83 & yes & no & no & yes & no & yes & j \\ \hline KIC\,9472174 & 12.3 & sdB+MV & 0.13 & -- & no & no & yes & no & yes & no & k \\ \hline \end{tabular} \label{tab:summary} \end{center} \flushleft References: (a) \citet{derekas11,borkovits12,fuller13}, (b) \citet{beck13}, (c) \citet{hekker10,frandsen13}, (d) \citet{southworth11,lehmann13}, (e) \citet{hambleton13}, (f) \citet{maceroni09}, (g) \citet{maceroni13} , (h) \citet{debosscher13}, (i) \citet{maceroni14}, (j) \citet{welsh11}, (k) \citet{ostensen10}. \end{small} \end{table*} Continued observations and the search for new asteroseismic eclipsing binaries remains crucial to answer important questions regarding stellar pulsations and evolution across the HRD. To highlight one particular aspect, Figure \ref{fig:scalingtest} shows updated empirical tests of the \numax\ and \Dnu\ scaling relations for stochastic oscillators. Note that most points in the \numax\ comparison are not fully independent from asteroseismology, with properties determined either by combining interferometric angular diameters with asteroseismic densities calculated from the \Dnu\ scaling relation \citep[e.g.,][]{huber12b}, or masses and radii determined from individual frequency modeling \citep[e.g.,][]{metcalfe14}. \ebgiant\ marks the first datapoint from an eclipsing binary, allowing an independent test of \numax\ and \Dnu\ that is otherwise only possible for wide binaries with radii and masses measured from astrometry and interferometry. The \Dnu\ comparison includes four stars which host exoplanets with independently constrained eccentricities, which allows an independent measurement of the mean stellar density \citep{seager03,winn10b}: HD17156 \citep{nutzman11,gilliland11}, TrES-2 \citep{barclay12,southworth11}, Hat-P7 \citep{cd10,southworth11}, and Kepler-14 \citep{huber13,southworth12}. Note that only stars with calculated \numax\ and \Dnu\ with uncertainties better than 20\% and 10\% have been included in the comparison. \begin{figure} \begin{center} \resizebox{8.3cm}{!}{\includegraphics{figures/testscaling.eps}} \caption{Empirical tests of asteroseismic scaling relations. Top: Comparison of measured \numax\ values to empirical values calculated from independent measurements (see legend). Filled symbols are empirical values which are independent of asteroseismology. Bottom: Same as top figure but for the \Dnu\ scaling relation.} \label{fig:scalingtest} \end{center} \end{figure} The median residuals for both quantities are close to zero, with a scatter of $7$\% for \numax\ and $3$\% for \Dnu. While these numbers are encouraging, it is important to note that the observational uncertainties for \numax\ and \Dnu\ derived from \kep\ data are typically up to a factor of 2 or more smaller \citep{chaplin13}. Additionally, comparisons for evolved giant stars are essentially limited to one datapoint, yet the vast majority of stars with asteroseismic detections are giants (see Figure \ref{fig:history}). Indeed, \ebgiant\ indicates that the \numax\ scaling relation remains accurate for giant stars, while the \Dnu\ scaling relation predicts a density which is too low by $\sim$7\%. Observations of stochastic oscillations in additional eclipsing binary systems will be required to investigate whether this offset may be systematic, and allow an \textit{empirical} calibration of scaling relations. Future observations of asteroseismic eclipsing binaries with giant and dwarf components can be expected from space-based missions such as K2 \citep{howell14}, TESS \citep{ricker09} and PLATO \citep{rauer13}. Importantly, these mission will observe stars which are significantly brighter than typical CoRoT and \kep\ targets, hence increasing the potential for independent constraints from ground-based observations such as long-baseline interferometry. There is little doubt that future space-based observations of asteroseismic eclipsing binary stars, combined with improved modeling efforts, will continue to play an important role to advance our understanding of stellar evolution across the H-R diagram. \vspace{0.5cm} \noindent \textbf{Acknowledgements:} I thank the organizers Elizabeth Griffin and Bob Stencel for a fantastic conference, and I am grateful to Paul Beck, J{\o}rgen Christensen-Dalsgaard, Orlagh Creevey, Jonas Debosscher, Aliz Derekas, Saskia Hekker, S{\o}ren Frandsen, Jim Fuller and Roy {\O}stensen for providing figures and comments on the manuscript. Financial support was provided by an appointment to the NASA Postdoctoral Program at Ames Research Center administered by Oak Ridge Associated Universities, and NASA Grant NNX14AB92G issued through the Kepler Participating Scientist Program. \newpage | 14 | 4 | 1404.7501 |
1404 | 1404.2324_arXiv.txt | We describe an open source GPU implementation of a hybrid symplectic $N$-body integrator, GENGA ({\bf G}ravitational {\bf EN}counters with {\bf G}pu {\bf A}cceleration), designed to integrate planet and planetesimal dynamics in the late stage of planet formation and stability analyses of planetary systems. GENGA uses a hybrid symplectic integrator to handle close encounters with very good energy conservation, which is essential in long-term planetary system integration. We extended the second order hybrid integration scheme to higher orders. The GENGA code supports three simulation modes: Integration of up to 2048 massive bodies, integration with up to a million test particles, or parallel integration of a large number of individual planetary systems. We compare the results of GENGA to Mercury and pkdgrav2 in respect of energy conservation and performance, and find that the energy conservation of GENGA is comparable to Mercury and around two orders of magnitude better than pkdgrav2. GENGA runs up to 30 times faster than Mercury and up to eight times faster than pkdgrav2. GENGA is written in CUDA C and runs on all NVIDIA GPUs with compute capability of at least 2.0. | The use of numerical $N$-body simulations to study the evolution of gravitational many-body systems, in particular that of our solar system, has a long tradition in astronomy. Prediction of planetary positions have occupied the field since the time of Newton, while at present the study of the full physical process of the formation and evolution of planetary systems requires a very significant amount of computing resources. We present here a new $N$-body integrator, GENGA\footnote{GENGA is available as open source code from https://bitbucket.org/sigrimm/genga} ({\bf G}ravitational {\bf EN}counters with {\bf G}pu {\bf A}cceleration), which uses today's most efficient computing hardware: the graphical processing units (GPUs). GENGA supports three computing modes: simulations of up to 2048 planetesimals, simulations with up to a million massless test particles, and parallel simulations of a large number of small planetary systems. \subsection{Physical Motivation} In the following we give a short overview of past work done in planetary $N$-body simulations which has motivated and guided the development of GENGA. This has mainly focused on four application areas, namely long-term evolution and stability analysis of the solar system, the dynamics of small asteroids under the gravitational influence of the planets, the planet formation process with the dynamics of planetesimals, and finally the evolution and characterization of exoplanetary systems. The long-term evolution and stability analysis of the solar system using $N$-body integrations has been studied by several people. A first result on an instability of the solar system was found by \cite{SussmannWisdom1988}, which integrated the five outer planets (including Pluto) over 845 Myr and found a Lyapunov time of Pluto's motion of 20 Myr. Including also the inner planets into the integration is more challenging because a much smaller time step is needed for a comparable accuracy. The entire solar system with all nine planets (including Pluto) and the Earth moon was performed by \cite{Quinn91} over 3 Myr backward in time. A longer simulation over 98.6 Myr was performed by \cite{SussmaWisdom92}, which used a symplectic $N$-body mapping \citep{WisdomHolman91, Gladman+1990, SahaTremaine1992} and confirmed the previous result on the chaotic motion of Pluto. They also confirmed the Lyapunov time of 5 Myr of the solar system, predicted by \cite{Laskar1989}. A small perturbation in the initial conditions of one of the inner planets can have dramatic effects on the evolution of the solar system. Even collisions between planets are possible in less than 3.5 Gyr \citep{Laskar96}. An overview of the question of stability of the solar system can be found in \cite{Laskar2012}. In addition to the massive planets in the solar system, massless test particles can be used to study the dynamics of meteorites, comets or impact ejecta. Since test particles do not interact with other test particles, the number of interactions that need to be calculated is greatly reduced. \cite{Gladman+1996} simulated 2100 particles escaping from Mars due to an impact and found a delivery efficiency to Earth of $7.5\%$ for $v_\infty = 1 \mathtt{kms^{-1}}$. They simulated also 200 particles escaping from Mercury and found one particle hitting the Earth after 23 Myr. The trajectories from Earth impact ejecta are studied by \cite{Wells+2003} with the PKDGRAV code \citep{Stadel2011} and they found 9 out of 675 particles returning to the Earth after 3000-5000 yr. They concluded that micro-biological ejecta could survive a sterilizing impact and reseed the Earth again with life. The delivery rates of terrestrial material to Mars and Venus and also back to Earth was studied by \cite{Gladman+2005}, and the trajectories of Mercurial material in more detail, by \cite{GladmanCoffey2009}. \cite{ReyesRuiz+2012} simulated $10^5$ test particles for 30,000 yr and found Earth ejecta reaching the Moon, Venus, Mars, Jupiter and Saturn, which means that biological material could in principle be transferred to other planets or their satellites by an impact to the Earth. To simulate the late stage of planet formation, starting from the runaway growth phase when planetesimals collide and form bigger objects like planetary embryos, test particles cannot be used anymore, and all $N^2$ gravitational interactions between the planetesimals have to be taken into account. Also close encounters between planetesimals can occur frequently and have to be resolved with a very small time step. These two effects already make the problem very challenging with a small number of bodies. Some of the first simulations of planetesimal dynamics in the inner solar system were able to integrate 100-200 bodies for $10^4-10^5$ yr \citep{ChambersWetherill1998,Aarseth+1993,Agnor+1999,Chambers2001}. The evolution of planetesimals and formation of planets was found to be a highly stochastic process and the solar system could not be reproduced well. A larger simulation containing $N\sim10^6$ planetesimals run for only 1000 yr by \cite{Richardson+2000} confirmed oligarchic growth, but was too short to study the entire planet formation process. The oligarchic growth was also studied in more detail over $4 \times 10^5$ yr with $10^4$ planetesimals by \cite{KokuboIda2002}. The process of terrestrial planet formation over more than 100 Myr was studied by simulating the dynamics of 1000-2000 planetesimals by, e.g., \cite{Raymond+06}, \cite{Obrien+06} and \cite{Morishima+2010}, by including an analytic gas disk model into the integration. These simulations can also be used to estimate the delivery rate of water or other volatile elements to the planets \citep{Elser+2012}. Since the full process of planet formation is stochastic, one cannot trust only one single simulation, but one has to study the statistics of many simulations with different initial conditions \citep{Kokubo+2006}. An overview of terrestrial planet formation with $N$-body simulations can be found in \cite{ChambersBook2011}. With the discovery of more and more exoplanetary systems with more than one planet, it was natural to study their dynamics and stability in a similar way to the work done on the solar system. Test particles can be used to find stable islands between detected exoplanets, which can help to predict additional planets, while long term simulations can be used to constrain the orbital parameters of the planets by analyzing the stability of the system. Many exoplanetary systems have been studied, e.g., by \cite{MenouTabachnik2003,Asghari+2004,RaymondBarnes2005}, and the here presented code was already used in \cite{Elser+13} to study the stability of hypothetical super Earths in the habitable zones of exoplanetary systems. \subsection{Technical Motivation} Since $N$-body simulations can require a large amount of computing power, it makes sense to use the fastest computer systems available, in order to save computing time. A review about used hardware in the history of $N$-body simulations can be found in \cite{BedorfZwart2012}. Two highlights in the history of special purpose computers for $N$-body simulations are the ``digital orrery'' \citep{Applegate+1985}, and the family of GRAPE (GRAvity PipE) computers \citep{Grape}. The digital orrery was a special machine built to integrate the equations of motion of planetary systems similar to the solar system. It consisted of a ring of processors, each one computing the trajectories of one planet. This machine was used to perform the integration from \cite{SussmannWisdom1988}. The GRAPE computers were able to compute the Newtonian force between two pairs of bodies directly in hardware. It was used as an accelerating device, sending the computed force between two particles to a central computer, on which the actual integration was performed. A short description of the different GRAPE types can be found in \cite{Grape}. Today's most efficient devices for $N$-body simulations are the GPUs. They consist of a large number of computing cores which can perform the same instructions on multiple threads (SIMT) in parallel. NVIDIA's GPUs can be programmed with the CUDA language (Compute Unified Device Architecture). Earlier methods to program GPUs are described in \cite{BedorfZwart2012}. First results of GPU-based general type $N$-body simulations were published by \cite{PortegiesZwart+2007}, \cite{Belleman+2007} or \cite{HamadaIitaka2007}. A modern implementation of the gravitational force calculation on GPUs, optimized for $N$ $>$ 1024, are described by \cite{Gems3} or \cite{CudaHandbook}. Codes using a Hermite integrator with block time steps for the general type $N$-body problem are given by the Sapporo library \citep{Sapporo}, the HiGPUs Code \citep{HiGPUs} or the NBODY6 Code \citep{NBODY6}. A library supporting the parallel integration of small $N$-body Keplerian systems is given by SWARM-NG \citep{SwarmNG}. \subsection{Contrast in Requirements with General $N$-body Simulations} The above listed codes are very efficient in solving the general type $N$-body problem, like in star clusters or cosmology simulations, where the individual bodies follow strongly non-Keplerian orbits. In planetary simulations by contrast, we can make the assumption that all bodies orbit a central mass following largely Keplerian arcs to lowest order with higher-order corrections resulting from mutual perturbations, where the solution of the Kepler problem can be computed analytically. Additionally, when a sufficient number of planetesimals are present within a system, close encounters may occur very frequently. The challenge is to conserve energy on secular timescales and yet treat close encounters properly, which is the focus of the present code. In order to demonstrate the importance of good energy conservation, we integrated the solar system using the HiGPUs code and compare the results with GENGA. In Figure\,\ref{fig:HiGPUs} is shown the evolution of the semi-major axes of the planets from the solar system. On the timescale of 500,000 yr, as shown in the Figure\,\ref{fig:HiGPUs}, the semi-major axes should remain almost constant, as reproduced by GENGA. In contrast the results of the HiGPUs code show a gradual but nonnegligible drift in the semi-major axis of the inner planets. Since the typical simulation for planet formation from planetesimals is typically integrated over 250 million yr, it makes it clear that a general $N$-body method should not be used for this problem. The HiGPUs code uses a high-order Hermite integrator with block time steps and integrates the general $N$-body problem without the assumption that the gravitational force of the central mass is dominant most of the time. To integrate the first 10,000 yr, HiGPUs needs about eight times more execution time than GENGA. For the rest of the simulation it needs around 50 times more time than GENGA. However, this is not to say that the HiGPUs code is inefficient in its typical application domain. In fact it is one of the most efficient codes, solving the classical gravitational $N$-body problem by using CUDA together with OpenMP and MPI. \begin{figure} \epsscale{1.25} \plotone{F1_color_low.eps} \caption{Evolution of the semi-major axes of the solar systems planets, integrated with the two codes GENGA and HiGPUs. While GENGA is designed exactly for planetary system integration and conserves the energy of the system very well, HiGPUs solves the general type $N$-body problem and the relative error in the energy is 50,000 times larger than with GENGA. As a consequence the inner planets begin to drift away from the Sun. To perform the shown integrations, HiGPUs needed about 50 times more execution time than GENGA.} \label{fig:HiGPUs} \end{figure} \subsection{Layout of the Paper} The structure of this paper is as follows. In Section\,\ref{symplectic} we describe the theory behind the GENGA integrator and the used numerical methods, followed by a description of the GPU in Section\,\ref{GPU} with some considerations that need to be made when writing parallel code for these devices. In Section\,\ref{GENGA kernels} we give an overview of the different kernels and a detailed description of the implementation. In Section\,\ref{Results} we compare GENGA with Mercury and pkdgrav2 regarding the energy conservation, a planet formation test and the performance. In Section\,\ref{Limits} we show and explain the limitations of the current code. | We presented the implementation of GENGA, a hybrid symplectic integrator designed and optimized for planetary system simulations. GENGA supports three simulation modes: Integration of up to 2048 massive bodies, integration with up to a million test particles, or parallel integration of a large number of individual planetary systems. We presented a detailed performance analysis of the code showing that at a large number of bodies, GENGA is up to 30 times faster than the Mercury code. At a very small number of bodies, GENGA is slower than Mercury due to GPU kernel overhead time and memory transfer between GPU and CPU. We compared the results of GENGA to Mercury and pkdgrav2 and found a very similar qualitative behavior of planetary systems between the codes. We showed that the energy conservation of GENGA is better than Mercury and much better that pkdgrav2. We presented the limitations of the current integration scheme and pointed out that future versions of GENGA should include an individual time stepping algorithm and a different changeover mechanism. GENGA expands the second order hybrid symplectic integration scheme to fourth and sixth order, and has successfully been used with the test particle and multi-simulation mode to analyze the stability of exoplanetary systems \citep{Elser+13}. GENGA is available as open source code from https://bitbucket.org/sigrimm/genga. | 14 | 4 | 1404.2324 |
1404 | 1404.5959_arXiv.txt | We introduce a mass dependent density profile to describe the distribution of dark matter within galaxies, which takes into account the stellar-to-halo mass dependence of the response of dark matter to baryonic processes. The study is based on the analysis of hydrodynamically simulated galaxies from dwarf to Milky Way mass, drawn from the MaGICC project, which have been shown to match a wide range of disk scaling relationships. We find that the best fit parameters of a generic double power-law density profile vary in a systematic manner that depends on the stellar-to-halo mass ratio of each galaxy. Thus, the quantity $\Mstar/\Mhalo$ constrains the inner ($\gamma$) and outer ($\beta$) slopes of dark matter density, and the sharpness of transition between the slopes ($\alpha$), reducing the number of free parameters of the model to two. Due to the tight relation between stellar mass and halo mass, either of these quantities is sufficient to describe the dark matter halo profile including the effects of baryons. The concentration of the haloes in the hydrodynamical simulations is consistent with N-body expectations up to Milky Way mass galaxies, at which mass the haloes become twice as concentrated as compared with pure dark matter runs. This mass dependent density profile can be directly applied to rotation curve data of observed galaxies and to semi analytic galaxy formation models as a significant improvement over the commonly used NFW profile. | \label{sec:introduction} Over several orders of magnitude in radius, dark matter (DM) halo density profiles arising from N-body simulations are well described by the so-called 'NFW' model \citep{Navarro96,Springel08,Navarro10}, albeit with well known systematic deviations \citep[e.g.,][]{Navarro04,Springel08, Gao08, Navarro10, DuttonMaccio2014}. The NFW function consists of two power laws, the inner region where the density is behaving as $\rho\propto r^{-1}$ and the outer part as $\rho\propto r^{-3}$. The central $\rho\propto r^{-1}$ ``cusps'' of such model disagree with observations of real galaxies where mass modeling based on rotation curves finds much shallower inner density slopes, known as ``cored'' profiles \citep[e.g.,][]{moore94,Salucci00,deBlok01,Simon05,Deblok08,Kuzio08,Kuzio09,oh11b}. Cored galaxies are also found within the fainter, dark matter dominated dwarfs spheroidal galaxies surrounding the Milky Way \citep{Walker11}. This {\it cusp/core discrepancy} is usually seen as one of the major problems of the $\Lambda$CDM paradigm at small scales. The NFW profile is, however, derived from pure DM simulations in which particles only interact through gravity. These simulations neglect hydrodynamical processes that may be relevant in determining the inner halo profile. Many studies have shown how baryons can affect the dark matter. Gas cooling to the center of a galaxy causes adiabatic contraction \citep[e.g.][]{Blumenthal86,Gnedin04}, whose effect strengthens cusps and exacerbates the mismatch between theoretical profiles and observations. Rather, expanded haloes are required to reconcile observed galaxy scaling relations of both early and late-type galaxies \citep{Dutton07,Dutton2013}. Baryons can expand haloes through two main mechanisms (see \citet{Pontzen14} for a recent review): outflows driven by stellar or AGN feedback \citep{Navarro96b,Mo04,Read05,Mashchenko06,Duffy10,Pontzen12,Martizzi13} and dynamical friction \citep{El-Zant01,Tonini06,Romano-Diaz08,DelPopolo09,DelPopolo10,Goerdt10,Cole11}. While dynamical friction is effective at expanding high mass haloes hosting galaxy clusters, stellar feedback is most effective at expanding low mass haloes \citep{governato10}. Gas cools into the galaxy centre where it forms stars that drive repeated energetic outflows. Such outflows move enough gas mass to create a core in an originally cuspy dark halo, due to the DM response to the adjusted gravitational potential. \citet{Penarrubia12} calculated the energy required to flatten a density profile as a function of halo mass. The cusp/core change can be made permanent if the outflows are sufficiently rapid \citep{Pontzen12}. Simulations from dwarf galaxies \citep{governato10,Zolotov12,teyssier13} to Milky Way mass \citep{Maccio12} have produced dark matter halo expansion depending on the implementation of stellar feedback. \citet{Governato12} showed that only simulated galaxies with stellar masses higher than $\sim 10^{7}M_\odot$ expand their haloes. They also showed that the inner DM profile slope, $\gamma$ in $\rho\propto r^{-\gamma}$, flattens with increasing stellar mass, resulting from the increase of available energy from supernovae. An increase in stellar mass may, however, also deepen the potential well in the central region of the halo: indeed, \citet{DiCintio13} showed that above a certain halo mass such a deepened potential well opposes the flattening process. \citet{DiCintio13} propose that $\gamma$ depends on the stellar-to-halo mass ratio of galaxies. At $\Mstar/\Mhalo\lesssim10^{-4}$ there is not enough supernova energy to efficiently change the DM distribution, and the halo retains the original NFW profile, $\gamma\sim-1$. At higher $\Mstar/\Mhalo$, $\gamma$ increases, with the maximum $\gamma$ (most cored galaxies) found when $\Mstar/\Mhalo$$\sim$$3-5\times10^{-3}$. The empirical relation between the stellar and halo mass of galaxies \citep{moster10,guo10} implies that this corresponds to $\Mstar$$\approx$$10^{8.5}M_\odot$ and $\Mhalo$$\approx$$10^{11}M_\odot$. In higher mass haloes, the outflow process becomes ineffective at flattening the inner DM density and the haloes have increasingly cuspy profiles. In this paper, we take the next step to provide a mass-dependent parametrization of the entire dark matter density profile within galaxies. Using high resolution numerical simulations of galaxies, performed with the smoothed-particle hydrodynamics (SPH) technique, we are able to study the response of DM haloes to baryonic processes. As with the central density slope $\gamma$ in \citet{DiCintio13}, we find that the density profile parameters depend on $\Mstar/\Mhalo$. This study is based on a suite of hydrodynamically simulated galaxies, drawn from the Making Galaxies In a Cosmological Context (MaGICC) project. The galaxies cover a broad mass range and include stellar feedback from supernovae, stellar winds and the energy from young, massive stars. The galaxies that use the fiducial parameters from \citet{stinson13} match the stellar-halo mass relation at $z=0$ \citep{moster10,guo10} and at higher redshift \citep{kannan13} as well as a range of present observed galaxy properties and scaling relations \citep{brook12b,stinson13}. Unlike previous generations of simulations, there is no catastrophic overcooling, no loss of angular momentum \citep{brook11,brook12a}, and the rotation curves do not have an inner peak, meaning that the mass profiles are appropriate for comparing to real galaxies. We present a profile that efficiently describes the distribution of dark matter within the SPH simulated galaxies, from dwarfs to Milky Way mass. The profile is fully constrained by the integrated star formation efficiency within each galaxy, $\Mstar/\Mhalo$, and the standard two additional free parameters, the scale radius $\rs$ and the scale density $\rhos$ that depend on individual halo formation histories. After converting $\rs$ into $r_{-2}$, i.e. the point where the logarithmic slope of the profile equals $-2$, we derive the concentration parameter for this new profile, defined as $c=\Rvir/r_{-2}$, and show that for high mass galaxies it substantially differs from expectation based on N-body simulations. This paper is organized as follows: the hydrodynamical simulations and feedback model are presented in \Sec{sec:simulation}, the main results, including the derivation of profile parameters and galaxies rotation curves, together with a comparison with N-body simulations in \Sec{sec:results} and the conclusions in \Sec{sec:conclusions}. | \label{sec:conclusions} It is well established that baryons affect dark matter density profiles of haloes in galaxies \citep[e.g.][]{Blumenthal86,Navarro96b,El-Zant01,Gnedin04,Read05,Goerdt06,Read06,Mashchenko06,Tonini06,Romano-Diaz08,DelPopolo09,governato10,Goerdt10,DiCintio11,Zolotov12,Governato12,Maccio12,Martizzi13,teyssier13}. Simple arguments compare the energy available from star formation with the depth of a galactic potential to estimate the degree of the change in the initial dark matter distribution \citep{Penarrubia12,Pontzen12,Pontzen14}. This study describes the dark matter profiles of haloes from a suite of hydrodynamical cosmological galaxy formation simulations that include the effects of stellar feedback. The profiles are modeled using a generic double power law function. We find that the slope parameters of such model ($\alpha,\beta,\gamma$) vary in a systematic manner as a function of the ratio between \Mstar/\Mhalo, which we call integrated star formation efficiency. Using these fits allows us to propose a star formation efficiency dependent density profile for dark matter haloes that can be used for modeling observed galaxies and in semi-analytic models of galaxy formation. The star formation efficiency dependent density profile has the form of a double power-law, with inner slope ($\gamma$), outer slope ($\beta$) and sharpness of transition ($\alpha$) fully determined by the stellar to halo mass ratio as given in Eq.~\ref{abg}. Thus, the five free parameters of the generic model reduce to two, the scale radius $\rs$ and scale density $\rhos$, the same free parameters of the commonly used NFW model. To examine how the scale radii varies as a function of integrated star formation efficiency, we compare the concentration parameter, $c=\Rvir/r_{-2}$, of the dark matter haloes from galaxies simulated with hydrodynamics prescriptions to those from the corresponding dark matter only simulations. For masses below roughly the Milky Way's the concentrations are similar, indicating that while the profiles may be significantly different from NFW, particularly in terms of inner slope, the radius at which the logarithmic slope of the profile equals -2 is the same as in the NFW model, indicating no net halo response at scales near the scale radius. However, for Milky Way mass galaxies the haloes from the hydro runs become as much as two times more concentrated than in the pure dark matter runs. Such high concentrations are consistent to what has been derived from observations of Milky Way's dynamical tracers \citep{Battaglia05,Catena10,Deason12a,Nesti13}. Thus, specifying the halo or stellar mass for a galaxy is sufficient to completely describe the shape of dark matter profiles for galaxies ranging in mass from dwarfs to L$^*$, based on the influence of stellar feedback. Importantly, the simulations we utilize in determining these profiles match a wide range of scaling relations \citet{brook12b}, meaning that their radial mass distributions are well constrained. \noindent The main features of the mass dependent dark matter profile are: \begin{itemize} \item Baryons affect the profile shape parameters. For galaxies with flat inner profiles $\gamma$ the sharpness of transition parameter, $\alpha$, increases from 1 to 3 and corresponds to a small decrease in the slope of the outer profile $\beta$. \vspace{0.1cm} \item At low integrated star formation efficiencies, $\Mstar/\Mhalo\lsim10^{-4}$ (galaxies with $\Mstar\lsim5$x$10^6\Msun$), dark matter haloes maintain the usual NFW profile as in dark matter only simulations. \vspace{0.1cm} \item At higher efficiencies the profile becomes progressively flatter. The most cored galaxies are found at $\Mstar/\Mhalo\approx3-5\times10^{-3}$ or \Mstar\ $\sim$$10^{8.5}$\Msun. \vspace{0.1cm} \item Galaxies with $\Mstar/\Mhalo\gtrsim5\times10^{-3}$ ($\Mstar\gtrsim10^{8.5}\Msun$), become progressively steeper in the inner region as their mass increases. \vspace{0.1cm} \item The parameters ($\alpha,\beta,\gamma$) returns to the NFW values of (1,3,1) for L$^*$ galaxies. \vspace{0.1cm} \item However such L$^*$ galaxies, and more in general galaxies with $\Mstar/\Mhalo\gtrsim0.03$, are up to a factor of 2.5 more concentrated than the corresponding dark matter only simulations. \end{itemize} In an Appendix we show step-by-step how to derive the dark matter profile for any galaxy mass. Our results show that baryonic effects substantially change the structure of cold dark matter haloes from those predicted from dissipationless simulations, and therefore must be taken into account in any model of galaxy formation. Of course, our model uses a particular feedback implementation, namely thermal feedback in the form of blast-wave formalism. Yet \cite{teyssier13} finds a similar degree of core creation, at least in low mass galaxies, using a different feedback scheme. Both studies are based on the same mechanisms for core creation, i.e. rapid and repeated outflows of gas which result in changes in the potential. Indeed, the simulations closely follow the analytic model of core creation presented in \citet{Pontzen12}, indicating that the precise details of the feedback implementation are not central to our results, at least not in a qualitative manner. Galaxy formation models which do not include impulsive supernova explosions driving outflows from the central regions will not form cores in this manner. In a forthcoming paper we will present a comprehensive comparison of our predicted density profile with the inferred mass distribution of observed galaxies, with particular emphasis on Local Group members. | 14 | 4 | 1404.5959 |
1404 | 1404.7392_arXiv.txt | The calcium-silicate-hydrate is used to model properties of cement on Earth. We study cementitious nanoparticles and propose these structures as components of cosmic dust grains. Quantum density functional theory methods are applied for the calculation of infrared spectra of Ca$_4$Si$_4$O$_{14}$H$_4$, Ca$_6$Si$_3$O$_{13}$H$_2$, and Ca$_{12}$Si$_6$O$_{26}$H$_4$ clusters. We find bands distributed over the near, mid and far-infrared region. A specific calcium-silicate-hydrate spectral feature at 14 $\mu$m, together with the bands at 10 and 18 $\mu$m which exist for other silicates as well, could be used for a detection of cosmic cement. We compare calculated bands with the 14 $\mu$m features in the spectra of HD 45677, HD 44179, and IRC+10420 which were observed by {\it Infrared Space Observatory} and classified as remaining. High abundance of oxygen atoms in cementitious nanoparticles could partially explain observed depletion of this element from the interstellar medium into dust grains. | \label{intro} Dust is produced in supernova explosions and the outflows of stars \citep{Draine2011,Tielens2013}. It is very important in the evolution of stars and galaxies. Dust grains act as a substrate for the formation of H$_2$ and many other astrophysical molecules, including prebiotic ones. It is accepted today that cosmic dust consists of silicates and carbonaceous materials with impurities from several chemical elements. Using infrared spectroscopy both silicates and carbon based grains have been observed in various environments: close to Earth and exoplanets, in the interstellar medium, in comets, around stars, and active galactic nuclei. The light from cosmic objects heats dust grains, which is followed by a dust emission in the infrared spectral region. Absorption features in the infrared produced by dust are also observed. The size of grains is from the length of one molecule, up to several hundred micrometers. It is known that 10 per cent of grains are ultra-small, with the size around 1.5 nm or less \citep{Li2001}. Cosmic dust mainly forms from the most abundant elements, such as silicon, carbon, oxygen, iron, and magnesium. Calcium is one of chemical elements that are produced in supernova explosions. It belongs to the group of twenty most abundant chemical elements in the Universe. The Ca II absorption was one of the first lines discovered in the interstellar space \citep{Hartmann1904}. It has been proposed that compounds which contain calcium exist in space. For example, spectra of Ca-poor and Ca-rich pyroxenes have been measured in the mid and far-infrared wavelengths in laboratories on Earth \citep{Koike2000}. It has been suggested that the amorphous diopside, CaMgSi$_2$O$_6$, contributes to the far-infrared spectrum of the planetary nebulae NGC 6302. Calcium has been measured in meteorites and rocks of Mars where it is a component in the calcium-aluminum-rich inclusions and chondrules \citep{Simon2009}. Calcium compounds are the main constituents of cement, which is one of the most used materials on Earth \citep{Allen2007,Pellenq2009,Skinner2010,Masoero2012}. The cement paste is formed when the cement powder, consisting mainly of alite (Ca$_3$SiO$_5$) and belite (Ca$_2$SiO$_4$), is mixed with water. The calcium-silicate-hydrate has a granular complex structure. Its chemical composition is variable, being characterized by the CaO:SiO$_2$ ratio in the range from 0.7 to 2.3 \citep{Manzano2007}. It has been found that temperatures above 350 K are necessary for the reaction between H$_2$O gas molecules and silicates in the outflows of stars \citep{Grossman1974}. However, silicate cosmic dust grains are often covered with water ice mantles. H$_2$O from mantles could react with silicates and calcium atoms and form a cosmic cement paste. We expect that cementitious materials could be produced as a dust component around oxygen-rich stars where silicates are dominant. The presence in space of various minerals in the form of hydrous silicates (for example, talc Mg$_3$[Si$_4$O$_{10}\mid$(OH)$_2$] and montmorillonite (Mg,Al)$_2$[Si$_4$O$_{10}\mid$(OH)$_2$](Na,K,Ca)$_x\cdot$nH$_2$O) has been under discussion for some time \citep{Whittet1997,Hofmeister2006,Mutschke2008}. Infrared spectroscopy is used in the field of solid state astrophysics \citep{Jager2011}, and astronomical observations are compared with spectra measured in laboratories. For example, spectra between 2.4 and 195 $\mu$m for 17 oxygen-rich circumstellar dust shells were observed using the Short and Long Wavelength Spectrographs on the {\it Infrared Space Observatory} ({\it ISO}) and compared with laboratory measurements \citep{Molster2002a,Molster2002b}. Although many bands were fitted with measured spectra of Mg-rich olivines and pyroxenes, 20 per cent of the spectral features were not identified. The {\it Spitzer} and {\it Herschel} missions collected many infrared spectra of dust grains \citep{Watson2009,Ciesla2014}. In laboratory measurements, in order to model various cosmic conditions, spectra of materials are studied from very low temperatures up to 1000 K. However, the interpretation of measured infrared spectra of complex materials can often be involved. Infrared spectroscopy of nanoparticles is even more demanding in comparison with measurements of the bulk material. Computational infrared studies of dust materials provide a connection between the microscopic structure and spectral properties. This is very important in astrophysics. For example, because of astronomical emissions at 3.3, 6.2, 7.7, 8.6, 11.2, and 12.7 $\mu $m, infrared spectra of polycyclic aromatic hydrocarbons (PAHs) have been calculated and assembled in a spectroscopic database \citep{Bauschlicher2010,Boersma2014}. Recently, spectra of PAHs clusters have been calculated \citep{Ricca2013}. Infrared spectra of a dust grain model in the form of the bare nanopyroxene cluster Mg$_4$Si$_4$O$_{12}$, as well as of its hydrogenated and oxygenated forms, have also been calculated by quantum computational methods \citep{Goumans2011}. Various crystalline and amorphous forms of silicates with different sizes, shapes and chemical compositions could form under diverse conditions in space \citep{Henning2010}. We have selected three clusters which consist of Ca, Si, O, and H atoms. Using these nanoparticles we model the crystalline and amorphous states of possible ultra-small cosmic dust grains with the structure of cement paste. Density functional theory computational methods \citep{Martin2004} are used to study structural properties of these nanoparticles, calculate vibrational modes, and explore important features in their infrared spectra. | \label{concl} Silicates are one of main components of cosmic dust grains. However, their precise chemical composition is not known. To the pool of possible silicate materials in cosmic dust we have added cement. We have studied one crystalline and two amorphous Ca-Si-H-O clusters which represent the chemical composition and bonding of cement at the nanoscale. Their infrared spectra are calculated, and it is found that bands are distributed over the entire infrared spectrum. We propose that the features at 14 $\mu$m, measured by the {\it Infrared Space Observatory} in the dust shells of several oxygen-rich stars and currently classified as unidentified \citep{Molster2002a,Molster2002b}, are the cement bands. From the abundance of Ca and other chemical elements we do not expect that cement nanoparticles are a dominant species in cosmic dust. However, cementitious nanoparticles, because of the many oxygen atoms they consist of, could act as an additional reservoir of O in cosmic dust and help in the solution of its depletion in the interstellar medium. With a recent detection of Ca-rich supernovae that produce much larger amounts of calcium then previously expected \citep{Perets2010,Kawabata2010}, as well as after discoveries of water everywhere in Space by the {\it Spitzer} and {\it Herschel} missions \citep{vanDishoeck2013}, all the necessary ingredients for the formation of cement nanoparticles in space are readily available. | 14 | 4 | 1404.7392 |
1404 | 1404.2781_arXiv.txt | { We used the PMAS integral field spectrograph to obtain large sets of radial velocities in the central regions of three northern Galactic globular clusters: \object{M3}, \object{M13}, and \object{M92}. By applying the novel technique of crowded field 3D spectroscopy, we measured radial velocities for about $80$ stars within the central $\sim10\arcsec$ of each cluster. These are by far the largest spectroscopic datasets obtained in the innermost parts of these clusters up to now. To obtain kinematical data across the whole extent of the clusters, we complement our data with measurements available in the literature. We combine our velocity measurements with surface brightness profiles to analyse the internal dynamics of each cluster using spherical Jeans models, and investigate whether our data provide evidence for an intermediate-mass black hole in any of the clusters. The surface brightness profiles reveal that all three clusters are consistent with a core profile, although shallow cusps cannot be excluded. We find that spherical Jeans models with a constant mass-to-light ratio provide a good overall representation of the kinematical data. A massive black hole is required in none of the three clusters to explain the observed kinematics. Our $1\sigma$ ($3\sigma$) upper limits are $5\,300\ \mathrm{M}_\odot$ ($12\,000\ \mathrm{M}_\odot$) for \object{M3}, $8\,600\ \mathrm{M_\odot}$ ($13\,000\ \mathrm{M}_\odot$) for \object{M13}, and $980\ \mathrm{M}_\odot$ ($2\,700\ \mathrm{M}_\odot$) for \object{M92}. A puzzling circumstance is the existence of several potential high velocity stars in \object{M3} and \object{M13}, as their presence can account for the majority of the discrepancies that we find in our mass limits compared to \object{M92}.} | \label{sec:introduction} Over the last years, the search for intermediate-mass black holes (IMBHs) has attracted remarkable attention. With $\sim10^2\text{--}10^4$ solar masses, they would bridge the gap from stellar-mass black holes to supermassive ones (SMBHs). Constraining their population statistics might answer the question how SMBHs assemble their masses. The scaling relations observed between SMBHs and fundamental properties of their host bulges, such as luminosities \citep{1995ARA&A..33..581K}, stellar masses \citep{1998AJ....115.2285M,2003ApJ...589L..21M,2004ApJ...604L..89H} or stellar velocity dispersions \citep{2000ApJ...539L...9F,2000ApJ...539L..13G,2009ApJ...698..198G} suggest that their growth is closely linked to the evolution of the host galaxy \citep[][but see also \citealt{2011ApJ...734...92J}]{1998A&A...331L...1S}. The progenitors of SMBHs are likely to be found in the building blocks of present-day galaxies. The close connection of Galactic globular clusters to the build-up of the Milky Way was proposed already by \citet{1978ApJ...225..357S} and is supported by cosmological simulations \citep[e.g.,][]{2006MNRAS.368..563M}. Globular clusters can therefore be considered as promising candidates to host IMBHs. Interestingly, a straightforward extrapolation of the SMBH scaling relations to the properties of globular clusters also yields black hole masses in the range $10^2\text{--}10^5$ solar masses. Clearly, such an extrapolation is a huge simplification, and there may be evidence that the common relations disagree with the observations already in the regime of low-mass galaxies \citep{2010ApJ...721...26G}. The runaway merging of massive stars in the early phases of cluster evolution has been suggested as a formation channel for IMBHs in dense star clusters \citep{2002ApJ...576..899P}. However, it has been argued \citep[e.g.][]{2009A&A...497..255G} that strong stellar winds restrict the resulting black holes to stellar masses ($\gtrsim10~\mathrm{M_\odot}$). An alternative formation scenario involves the collapse of massive population III stars \citep[e.g.][]{2001ApJ...551L..27M}. Observational evidence for (but also against) the existence of IMBHs is still extremely scant. Gas accretion onto such a black hole would allow for its detection using radio or X-ray observations, such as in the case of the IMBH candidate HLX-1 in the galaxy \object{ESO~243-49} \citep{2009Natur.460...73F}. \object{G1} in \object{M31}, the most massive known globular cluster in the local group, was detected in both radio \citep{2007ApJ...661L.151U} and X-ray observations \citep{2004ApJ...616..821T,2010MNRAS.407L..84K}, although \citet{2012ApJ...755L...1M} could not confirm the radio detection. Recent observations by \citet{2012ApJ...750L..27S} place stringent upper limits on the amount of radio emission coming from Galactic globular clusters. Translating those into mass limits, however, requires making assumptions about the accretion physics that are not well understood. Several authors investigated the possibility that IMBHs imprint their presence onto photometrically observable properties of a globular cluster. \citet{2005ApJ...620..238B} and \citet{2011ApJ...743...52N} found that a massive black hole should produce a shallow cusp in the central surface brightness profile of the surrounding cluster. A large ratio of core to half-mass radius was suggested as indirect evidence for the presence of an IMBH by \citet{2007MNRAS.374..857T}. Furthermore, the existence of extreme horizontal branch stars has been proposed as a tracer for IMBHs by \citet{2007MNRAS.381..103M}. \citet{2008ApJ...686..303G} investigated the influence of black holes on mass segregation among the cluster stars. Arguably, the most direct way to not only find massive black holes but also obtain their masses is the detection of their kinematic fingerprints. The best SMBH mass estimates were obtained from stellar kinematics in our own galaxy \citep{2009ApJ...692.1075G} or gas kinematics in \object{NGC~4258} \citep{2005ApJ...629..719H}. However, obtaining meaningful kinematic measurements in the central regions of globular clusters is a challenging task: While the measurement of individual stellar velocities is hampered by crowding, integrated-light analyses can be significantly affected by shot noise from the few brightest giants \citep{1997A&A...324..505D}. Kinematic studies of the centres of globular clusters remained ambiguous so far, in some cases even contradictory conclusions were reached for the same clusters. A prominent example is \object{$\omega$~Centauri}, for which \citet{2010ApJ...710.1032A} and \citet{2010ApJ...710.1063V} find no evidence for an IMBH while \citet{2010ApJ...719L..60N} and \citet{2012A&A...538A..19J} claim the detection of one, with a mass of $4\times10^4\ \mathrm{M_\odot}$. Similarly, the detection of an IMBH with $17\,000$ solar masses in \object{NGC~6388} by \citet{2011A&A...533A..36L} was not confirmed by \citet{2013ApJ...769..107L}, who obtain an upper limit of $2\,000\ \mathrm{M_\odot}$ instead. Further detections of IMBHs have been reported in a small number of massive clusters, \object{G1} \citep{2005ApJ...634.1093G} among them. In a sample recently studied by \citet{2012A&A...542A.129L,2013A&A...552A..49L}, the kinematics in $2$ out of $7$ clusters suggested the presence of an IMBH, while for the remaining clusters upper limits of typically $>1\,000$ solar masses were derived. Similar mass limits were also reported in other studies, carried out for \object{M15} by \citet{2002AJ....124.3270G} and \citet{2006ApJ...641..852V}, for \object{47~Tuc} by \citet{2006ApJS..166..249M} or for \object{NGC~6266} by \citet{2012ApJ...745..175M}. Clearly, a conclusive picture of which globular clusters host IMBHs is not established yet. Consequently, the question whether the scaling relations established for SMBHs can be extrapolated into the regime of globular clusters is also unanswered. So far the results are still consistent with it. Alternatively, IMBHs might follow different scaling relations, as suggested by \citet{2007MNRAS.381..103M}. More observations are the only way to make progress here. So far, all claimed IMBH detections come from integrated light spectroscopy, whereas studies based on the kinematics of resolved stars derived upper limits which in some cases are in conflict with the detections from the former approach. This may suggest that the influence of the few brightest stars still hampers the integrated light measurements. In any case, resolving this situation requires new techniques for spectroscopy in crowded stellar fields and a better understanding of its capabilities and limitations. In \citet[][hereafter \citetalias{2013A&A...549A..71K}]{2013A&A...549A..71K}, we recently presented a new method for analysing integral field spectroscopy (IFS) data of such fields. It extends the established analysis techniques for crowded field photometry into the domain of three-dimensional datacubes by fitting a wavelength dependent point spread function (PSF) to deblend stellar spectra. In this paper, we apply this technique to IFS data of three Galactic globular clusters, \object{M3}, \object{M13}, and \object{M92}, with the aim of constraining the presence of intermediate-mass black holes in the objects. The paper is organized as follows. After laying out the the target selection, the observations, and their reduction in Sects.~\ref{sec:sample} to \ref{sec:reduction}, we describe the analysis of the photometric (Sect.~\ref{sec:photometry}) and the kinematic data (Sect.~\ref{sec:kinematic}). In Sect.~\ref{sec:dynamics}, the cluster dynamics are analysed and the search for intermediate-mass black holes in the clusters via Jeans modelling is performed. Our findings are discussed in Sect.~\ref{sec:discussion}, and we provide our conclusions in Sect.~\ref{sec:conclusions}. | \label{sec:conclusions} We introduce crowded field 3D spectroscopy as a powerful tool to study the central kinematics of Galactic globular clusters. The combination of PMAS integral field data with PSF fitting techniques allows us to gather large samples of stellar spectra in the highly crowded centres of the three clusters \object{M3}, \object{M13}, and \object{M92}. In our seeing limited datasets, individual stars are resolvable down to the main sequence turn-off of each cluster. As expected, the S/N of the deblended spectra decreases as the stars get fainter, yet reliable radial velocity measurements are still possible for stars below the horizontal branch. The combination of our new measurements with literature data allows us to gather kinematic information at all cluster radii. All our catalogues of radial velocity measurements are made available in the online version of this paper. Our PMAS stars constitute by far the largest spectroscopic samples obtained in the central $10$--$20\arcsec$ of any of the three clusters to date. They allow us for the first time to put meaningful constraints on the presence of intermediate-mass black holes in these clusters. Using spherical Jeans models, we find that a massive black hole is required in none of them to reach agreement between the Jeans models and our data. Our analysis thus puts stringent upper limits on the mass of any such object in the investigated clusters. In \object{M92} we can rule out an intermediate-mass black hole more massive than $2\,700\ \mathrm{M_\odot}$ at the $3\sigma$ level. Our $3\sigma$ upper limits for \object{M3} ($12\,000\ \mathrm{M_\odot}$) and \object{M13} ($13\,000\ \mathrm{M_\odot}$) are more relaxed. \object{M13} remains the least conclusive object we investigated. In contrast to the other two clusters, the most likely model for this cluster does contain a black hole (of $2\,000\ \mathrm{M_\odot}$). However, the improvement over the model without a black hole is insignificant. In this respect, it is unfortunate that the kinematical data in the centre of \object{M13} suffered from non-optimal observing conditions. In summary, we consider \object{M13} the most promising candidate to potentially host an IMBH, still below our sensitivity limit but possibly detectable with better data. The analysis of the central kinematics in the two clusters \object{M3} and \object{M13} is complicated by the presence of a few potential high velocity stars. Excluding these stars from the analysis would strongly narrow the range of black hole masses allowed by our modelling. Our current dataset is insufficient to discriminate whether these stars represent the wings of the internal velocity dispersion or have been accelerated by other means. Acceleration by an undetected intermediate-mass black hole is unlikely in view of the relatively large distances to the centres and the sizes of the stellar samples \citep{2003ApJ...597L.125D}. As mentioned earlier, another possible scenario involves the acceleration by stellar mass black holes. Heating by stellar mass black holes has also been suggested as an explanation for large core radii \citep{2008MNRAS.386...65M,2012Natur.490...71S}. In this respect, the large core radii of \object{M3} \citep[$1.1\mathrm{pc}$, ][]{2005ApJS..161..304M} and \object{M13} ($1.2\mathrm{pc}$) are intriguing, they may suggest a wealthy black hole population in the centres of the clusters and larger kinematic samples will be able to investigate this. The search for intermediate-mass black holes in globular clusters has gained a lot of attention over the last years. Many of the recent studies rely on observations at the highest possible spatial resolutions. In the future such observations may be used to address the questions beyond the sensitivity of our dataset. However, the requirement for adaptive optics or multi-epoch HST observations makes them rather expensive in terms of observing time. We demonstrated in this work that meaningful results can be obtained based on less demanding observations. It is also clear that significant further improvements can be expected from applying the technique of crowded field 3D spectroscopy to higher quality data. | 14 | 4 | 1404.2781 |
1404 | 1404.2871_arXiv.txt | {The net effect of the small scale magnetic field on the Sun's (bolometric) brightness is studied with realistic 3D MHD simulations. The direct effect of brightening within the magnetic field itself is consistent with measurements in high-resolution observations. The high `photometric accuracy' of the simulations, however, reveal compensating brightness effects that are hard to detect observationally. The influence of magnetic concentrations on the surrounding nonmagnetic convective flows (a `proximity effect') reduces the brightness by an amount exceeding the brightening by the magnetic concentrations themselves. The net photospheric effect of the small scale field ($\approx-0.34\%$ at a mean flux density of 50 G) is thus negative. We conclude that the main contribution to the observed positive correlation between the magnetic field and total solar irradiance must be magnetic dissipation in layers around the temperature minimum and above (not included in the simulations). This agrees with existing inferences from observations. | The effect of the small scale magnetic field on (bolometric) brightness appears to have three distinct components: the brightness of the magnetic structure itself (composed of the bright interior of the structure at disk center, and the`bright wall effect' towards the limb), plus the two proximity effects it has on its surroundings: the `dark ring' resulting from the influx of radiation into the magnetic concentration, and the interference of magnetic concentrations with the nearby convective flow. The most conspicuous component is the bright wall effect, easily measurable as a brightening in active regions when seen near the solar limb. It has also been reproduced convincingly in realistic 3-D MHD simulations such as Carlsson et al.\ (2004), De Pontieu et al.\ 2006, and the present ones (cf.\ Fig.\ \ref{compo}). The $0\farcs1-0\farcs2$ narrow dark rings are also conspicuous in high resolution continuum images near disk center, but are less easily quantifiable because of the variable shapes of the structures (`crinkles'). Finally, the effect on the surrounding convective flow pattern is well known from observations, but its effect on brightness is hard to detect, smeared out over too large an area to be measurable at the photometric accuracy of ground-based observations. It has been detected however, in data from Hinode (Kobel et al. 2012). \begin{figure}[t]% \center\includegraphics[width=0.9\hsize]{comm_plots/velocity.pdf} \center\includegraphics[width=0.9\hsize]{comm_plots/velocitydiff.pdf} \caption{Top: average vertical velocity $\langle v_z\rangle$ as a function of distance from magnetic elements (solid) compared with $\langle v_z\rangle$ near flow convergence points of the nonmagnetic simulation (broken). Bottom: difference between the two, showing the proximity effects on vertical velocities.}\label{velocity} \end{figure} Perhaps surprisingly, the negative contribution of the proximity effects turns out to dominate the photospheric brightness change. The net brightness effect is thus the opposite of the standard prediction (S77). Observationally, the effect of magnetic concentrations on nearby convective flows is easily detectable through changes in granulation morphology and vertical velocities (`anomalous granulation', cf. Title et al. 1986, 1992, Kobel et al. 2012). Concerns that these changes could also affect energy transport and hence surface flux have been around for some time (e.g.,\ Spruit 1998). They were not discussed much, possibly because the effect was not large enough to be detected with ground-based photometric accuracy. Our simulations also show changes in vertical velocities near magnetic concentrations. The spatial coincidence of these changes with changes in bolometric flux support the interpretation that the net darkening is caused by interference with the convective heat flux near magnetic concentrations. The effect appears to take place in a rather narrow region, extending from the intergranular lane to somewhat into the surrounding granule. The magnetic concentrations are larger on average in the 100 G simulation than at 50 G. Their effect on the surrounding flows is correspondingly somewhat larger (Fig.\ \ref{velocity} bottom panel). The net negative effect on bolometric brightness does not differ much in the 50 G simulation. We interpret this as a consequence of the compensating bright wall effect. Its increase towards the limb is most prominent in larger concentrations that are less affected by self-obscuration away from disk center. This is evident in the disproportionately larger brightness increase of the 100 G simulation towards the limb compared with the 50 G result (Fig.\ \ref{clv}). It can also be seen qualitatively in the CLV of the images in Fig.\ \ref{compo}. This leaves the question how the observed positive correlation of total solar irradiance (TSI) with the small scale field comes about. The most likely explanation is that the simulations underestimate the contribution of chromosphere and upper photosphere. A major contribution of the chromosphere to TSI has in fact already been inferred from the wavelength dependence of solar irradiance variability. Unruh et al. (1999) concluded that the photosphere contributes negligibly to TSI variation, compared with the chromosphere. More recently, Ball et al. (2012) estimate the photospheric contribution at 18\%. Empirical models for the mean stratification in active regions such as Vernazza et al. (1973) already indicated the presence of heating processes starting around the temperature minimum. This has been interpreted as evidence of some form of magnetic heating. Our calculations necessarily miss most of this contribution because of the use of a potential field as upper boundary condition. This forces the field near the upper boundary to its lowest energy state, from which no energy can be extracted. Proper inclusion of magnetic dissipation in simulations like the present ones is very demanding, because of the time step limitations resulting from the high Alfv\'en speed in the chromosphere. Simulations with methods adapted to this situation may be needed, such as have been developed by Gudiksen and Nordlund (2005) for the coronal heating problem. \\ | 14 | 4 | 1404.2871 |
|
1404 | 1404.7453_arXiv.txt | Ground-based ultra-high contrast imaging, as required for direct imaging of exoplanets and other solar systems, is limited by difficulty of separating the planetary emission from the effects of optical aberrations that are not compensated by the adaptive optics (AO) system, so-called ``non-common path aberrations" (NCPAs). Simultaneous ($\sim$ millisecond) exposures by the science camera and the AO system enable the use of ``phase diversity" to estimate both the NCPAs and the scene via a processing procedure first described by the author (R. Frazin 2013, ApJ, 767, article id. 21). This method is fully compatible with more standard concepts used in long-exposure high-contrast imaging, such as angular differential imaging and spectral deconvolution. Long-exposure methods find time-dependent NCPAs, such as those caused by vibrations, particularly challenging. Here, an NCPA of the form of $\alpha \cos(k \cdot r - \omega t + \vartheta)$\ is considered. It is shown that, when sampled at millisecond time-scales, the image plane data are sensitive to $\mbox{arg}(\alpha)$, $\vartheta$\ and $\omega$, and, therefore such NCPAs can be simultaneously estimated with the scene. Simulations of observations with ms exposure times are reported. These simulations include substantial detector noise and a sinusoidal NCPA that places a speckle exactly at the location of a planet. Simulations show that the effects of detector noise can be mitigated by mixing exposures of various lengths, allowing estimation of the planet's brightness. | \begin{wrapfigure}{r}{0pt} \begin{tabular}{rr} \epsfig{file=fig_abmod_a.eps,width=.3\linewidth,clip=} & \epsfig{file=fig_abmod_b.eps,width=.3\linewidth,clip=} \\ \epsfig{file=fig_abmod_c.eps,width=.3\linewidth,clip=} & \epsfig{file=fig_abmod_d.eps,width=.3\linewidth,clip=} \end{tabular} \caption{\small Stellar coronagraph simulation of a single ms science camera image showing how the instantaneous AO residual interacts with an NCPA of the form $\alpha \cos( k \cdot r + \vartheta)$. Each panel corresponds to an NCPA with the same $k$ and $|\alpha|$, but differing in spatial phase $\vartheta$ and whether $\alpha$\ is real or imaginary. This value of $k$\ places a speckles at the locations indicated by the arrows, at a distance of $4 \lambda/D$\ from the center. In the \emph{top left} image, $\alpha$\ is real and $\vartheta=0$. Similarly, the \emph{top right}, \emph{bottom left} and \emph{bottom right} correspond to $\alpha$\ values that are real, imaginary and imaginary, and $\vartheta$\ values of $\pi/2$, $0$ and $\pi/2$, respectively. This shows that the science camera data at ms cadence are sensitive to $\mbox{arg}(\alpha)$ and $\vartheta$. The $x-$\ and $y-$\ axes are in pixel units. In all panels, the intensity, with units of photons/ms, is displayed. All panels have the same color scale.} \label{AbMod} \vspace{-2mm} \end{wrapfigure} \vspace{-1.8mm} The idea of using millisecond exposures for ultra-high contrast imaging, such as is required for exoplanet science, has been given little attention, mostly due to the fact that all of the current high-contrast efforts are based on long exposures. The power of millisecond exposure analysis comes from two facts: \emph{1) most of the atmospheric motion is nearly frozen on that time-scale and 2) the wavefront sensor (WFS) provides a great deal of information about the AO residual}. The key point is that at every millisecond the AO residual presents new random wavefronts. When combined with knowledge of the wavefront, each millisecond exposure provides more information about the NCPAs. Using the WFS and millisecond images from the science camera has been dubbed ``random phase diversity," and it has been explored by [\citenum{Frazin13}] (henceforth, Paper I) and [\citenum{Codona13}]. In long-exposure imaging, far less information is available, as one only sees the average of all of the speckles, making the data from the WFS of little utility. Showing a stellar coronagraph simulation result from Paper I, Fig.~\ref{fig_modulation} shows time series of the intensity seen in a single pixel of the science camera. The dotted curve illustrates how the AO residual modulates the stellar speckle (which, for this pixel, is enhanced by a sinusoidal NCPA) at a cadence of 1 ms. The solid curve shows the much weaker modulation of the planetary light in the same pixel. These two time-series, the planetary intensity and the speckle intensity, are quite different in character, with the speckle having an approximately exponential probability density function (PDF), while the PDF of the planetary intensity is somewhat localized around its non-zero mean.\cite{Gladysz10,Frazin13} This can be understood as follows: The planetary time-series is stabilized by the AO system, as the flat part of the planet's wavefront is responsible for its intensity at this position in the image plane. However, this star's speckle is entirely due to the random, non-flat part of the star's wavefront (the coronagraph removes the flat part) and, hence, it is much more volatile. Long-exposure observations have difficulty distinguishing amplitude NCPAs from phase NCPAs. Consider pupil plane NCPA (upstream of the coronagraph) of the form $\phi(r) = \alpha \cos(k \cdot r + \vartheta)$, where $r$\ is pupil plane coordinate vector, $k$\ is the vector spatial frequency of the aberration, $\vartheta$\ is the spatial phase, and $\alpha$\ is the complex amplitude of the aberration.\footnote{Thus, the pre-coronagraph, pupil plane field is modified by this NCPA via multiplication: $E(r) \exp[ j \phi ]$, where $E(r)$\ would be the field without this NCPA. } The imaginary part of the NCPA corresponds to an error in the wavefront amplitude and the real part corresponds to a phase error. However, millisecond observations clearly separate these effects, as is illustrated in Fig.~\ref{AbMod}, which simulates four different sinusoidal NCPAs all using the same AO residual as input. These four NCPAs all have same spatial frequency $k$, which places the two indicated speckles at a distance of about $4\lambda/D$\ from the center, were $\lambda$\ is the wavelength. The top panels correspond to real NCPAs ($\alpha$\ is purely real), differing only in the phase angle $\vartheta = (0,\pi/2)$. The bottom panels correspond to purely imaginary NCPAs ($\alpha$\ is purely imaginary), again differing only in the phase angle $\varphi = (0,\pi/2)$. Of course, \emph{if the AO residual were precisely zero, all four of these NCPAs would lead to exactly the same two identical speckles.} Standard long-exposure images cannot distinguish between these aberrations, although one could solve for them using well-calibrated offsets of the AO system's deformable mirror,\cite{Thomas10} in a manner similar to the procedures given in Paper I. \begin{wrapfigure}{l}{0pt} \epsfig{file=fig_vibrations.eps,width=.4\linewidth,clip=} \caption{\small Stellar coronagraph simulation of the effect of a vibration on a speckle in the science camera. This speckle is caused by a vibrational NCPA of the form $\cos(k \cdot r - \omega t)$, and the modulation is caused by the interaction of the NCPA with the random AO residual. The black dotted curve corresponds to the intensity of a speckle for $\omega/2\pi =100$ Hz and the red solid curve corresponds to $\omega/2\pi = 10$Hz. Thus, the ms data is clearly sensitive to the frequency of the vibration. \vspace{-15mm}} \label{Vibration} \end{wrapfigure} Vibrations pose a particularly challenging problem to high-contrast astronomy. Since they are sensitive to specific conditions such as wind and thermal state of the system, they are difficult to characterize with reference images. Consider a vibration giving rise to an NCPA given by the form $\phi(r) = \alpha \cos(k \cdot r + \omega t)$, where $\omega$\ is the vibration frequency and $t$ is the time. Clearly, at millisecond time-scales, this NCPA will manifest a modulation in the science camera that depends on $\omega$. Fig.~\ref{Vibration} shows stellar coronagraph simulations similar to those in Paper I demonstrating that the intensity time-series of the speckle associated with this NCPA is indeed strongly dependent on $\omega$. | This paper has discussed millisecond cadence exposures in the science camera combined with simultaneous data from the wavefront sensor for the purpose of simultaneous exoplanet imaging and estimation of NCPAs. This mode of operation has the useful potential to detect both amplitude and phase NCPAs as well as time-dependent NCPAs caused by vibrations [see Figs.~\ref{AbMod} and \ref{Vibration}]. The amplitudes of these aberrations can be estimated in the framework given in Paper I.\cite{Frazin13} In addition, this paper gives results for simulations that include the effects of detector noise and shows that observation sequences that combine various exposure lengths can mitigate its effects. Millisecond imaging in the science camera may have the potential to revolutionize ground-based exoplanet imaging, but a number practical issues need to be explored first. These include the effects of uncertainties in wavefront sensor measurements, detector noise, chromatic effects, and others, some of which are listed in Paper I. It is important to recognize that short-exposure analysis is fully compatible with ADI and SD, as well as methods that require additional hardware to achieve phase diversity, since they can all be used in short-exposure mode. Short exposures are also compatible with (and probably useful for) ``electric field conjugation," a technique in which the deformable mirror (DM) is adjusted in order minimize the intensity in a specific region of the image plane, thereby increasing the contrast.\cite{Thomas10} Indeed, NCPAs determined via reference images or other types of long exposure analysis could be used a statistical priors on the estimation of the NCPAs with millisecond exposure data. Thus, many current efforts are quite applicable to the millisecond exposure paradigm. | 14 | 4 | 1404.7453 |
1404 | 1404.0335_arXiv.txt | We present high angular resolution ($\sim$0.5$^\prime$$^\prime$) MIR spectra of the powerful radio galaxy, Cygnus A, obtained with the {\it Subaru} telescope. The overall shape of the spectra agree with previous high angular resolution MIR observations, as well as previous {\it Spitzer} spectra. Our spectra, both on and off nucleus, show a deep silicate absorption feature. The absorption feature can be modeled with a blackbody obscured by cold dust or a clumpy torus. The deep silicate feature is best fit by a simple model of a screened blackbody, suggesting foreground absorption plays a significant, if not dominant role, in shaping the spectrum of Cygnus A. This foreground absorption prevents a clear view of the central engine and surrounding torus, making it difficult to quantify the extent the torus attributes to the obscuration of the central engine, but does not eliminate the need for a torus in Cygnus A. | Active galactic nuclei (AGN) include many different object classes with diverse properties. They span as many as eight decades in luminosity, include objects with and without broad emission lines, and have a wide variety of spectral energy distributions (SED). The torus, proposed in the unified models of AGN (see e.g., \citealt{1993ARA&A..31..473A,1995PASP..107..803U} for reviews of the models), allows for different observed classes of objects to be explained by the viewing angle of the observer to the AGN, instead of physical differences between the classes. In order to determine whether the difference in class is due to the viewing angle to the torus or an actual physical difference in the object, comparison of the properties of the torus for each class is needed. The size, structure, and geometry of the torus are not well-constrained. Early modeling work proposed a uniform dust density for reasons of computational tractability \citep{1988ApJ...329..702K,1992ApJ...401...99P,1993ApJ...418..673P,1994MNRAS.268..235G,1995MNRAS.273..649E}. These models produced geometrically-thick tori a few hundred parsecs in radius \citep{1994MNRAS.268..235G,1995MNRAS.277.1134E}. Since the energy that is absorbed from the central engine must be re-emitted at mid-infrared (MIR) wavelengths, the models were first tested against early {\it Infrared Astronomical Satellite} ({\it IRAS}) data and other far-infrared (FIR) data sets, with which they were consistent \citep{1995MNRAS.277.1134E,1997ApJ...486..147G}. However, high angular resolution ($\leq$0.5$^\prime$$^\prime$) MIR observations of AGN on 8m class ground-based telescopes require much more compact tori, such as found for NGC4151 (upper limit on MIR size of $\lesssim$35 pc, \citealt{2003ApJ...587..117R}), Circinus (upper limit on MIR size of 12 pc, \citealt{2005ApJ...618L..17P}), NGC1068 (upper limit of 15 pc, \citealt{2006ApJ...640..612M}), Centaurus A (upper limit on MIR size of 3.5 pc, \citealt{2008ApJ...681..141R}), and M87 \citep{2001ApJ...561L..51P,2007ApJ...663..808P}. More recently, MIR interferometry has found dust components at these scales for NGC1068 \citep{2004Natur.429...47J}, Centaurus A \citep{2007A&A...471..453M}, Circinus \citep{2007A&A...474..837T} and NGC4151 \citep{2009ApJ...705L..53B}. Larger samples of AGN observed with MIR interferometry also confirm compact tori for most objects in the samples \citep{2009A&A...502...67T,2013A&A...558A.149B}. Clumpy torus models fit the observed spectra and produce a torus scale consistent with the ground-based observations \citep{2006ApJ...640..612M,2009ApJ...693L.136M,2009ApJ...702.1127R,2009ApJ...707.1550N,2011ApJ...731...92R,2008ApJ...675..960P,2008A&A...485...33H,2010A&A...515A..23H} High resolution MIR imaging and spectroscopy has revealed important details of AGN, as seen in the subarcsecond variation detected for NGC1068 \citep{2006ApJ...640..612M} and Circinus \citep{2006MNRAS.367.1689R}. According to the unified scheme there should be little difference in the torus structure between radio quiet (RQ) and radio loud (RL) AGN, except for the presence of radio jets. To date, nearly all the high angular resolution observations of AGN have been done on RQ objects, which comprise 80-90\% of AGN. The first high angular resolution MIR imaging survey including a significant number of RL AGN was presented by \citet{2010A&A...511A..64V}. We have also recently completed high angular resolution imaging observations of six more RL AGN. \citet{2012AJ....144...11M} presents new imaging of four more RL AGN, and several more are presented in \citet{2011A&A...536A..36A}. High resolution imagine and spectroscopy has, up to now, only been done for 2 RL AGN: M87 (\citealt{2001ApJ...561L..51P} (imaging), \citealt{2007ApJ...663..808P} (spectroscopy)) and Centaurus A (\citealt{2007A&A...471..453M} (spectroscopy), \citealt{2008ApJ...681..141R} (imaging)). \begin{figure}[t] \centerline{\includegraphics[scale=.5]{f1.eps}} \caption{Left: N band ($\lambda$$_o$$=$10.8 $\mu$m, $\Delta$$\lambda$$=$5.2 $\mu$m) image of Cyg A with bicones (parabolas) and radio axis (dashed line) overplotted (Figure 4, \citealt{2002ApJ...566..675R}). Right: 2.25 $\mu$m NICMOS image of Cyg A from \citet{1999ApJ...512L..91T}. These images show the unresolved nucleus and surrounding extended emission. Overplotted on the images is the slit and the extraction apertures used for our observations. Region 1 is the southeastern aperture, region 2 is the central aperture, and region 3 is the northwestern aperture. The total aperture is the combination of all three regions. See $\S$ \ref{sec:subobs}.} \label{fig:slit} \end{figure} Here we present high angular resolution MIR spectroscopy for Cygnus A (Cyg A). The redshift of Cyg A, z=0.056 \citep{1994AJ....108..414S}, corresponding to a distance of 247 Mpc and a scale of 1$^\prime$$^\prime$=1.1 kpc, makes it one of the closest powerful FR II radio galaxies. Because of its proximity and brightness, Cyg A is very well-studied. Its radio jet structure and overall radio structure \citep{1998A&A...329..873K,1999AJ....118.2581C} have provided critical data for models of jets. X-ray spectroscopy found an intrinsic luminosity of the central core of $\sim$10$^{44}$ erg s$^{-1}$ and a large H\,\textsc{i} column density ($\sim$10$^{23}$ cm$^{-2}$) \citep{2002ApJ...564..176Y}. Optical spectropolarimetry \citep{1997ApJ...482L..37O} also found evidence of a powerful, obscured central engine. Furthermore, the polarization of narrow lines and the ratio of broad lines are consistent with scattering by dust. {\it Chandra} observations \citep{2002ApJ...564..176Y} found evidence of a heavily-obscured nucleus and a biconical soft x-ray emission that is aligned with the bipolar cone seen in optical emission lines \citep{1998MNRAS.301..131J} and near-infrared (NIR) \citep{1999ApJ...512L..91T,2000MNRAS.313L..52T} observations. High resolution MIR imaging observations (\citealt{2002ApJ...566..675R}, left panel of Figure \ref{fig:slit}) show an unresolved nucleus surrounded by extended emission that has a morphology consistent with the bipolar structure seen in optical and low resolution NIR emission lines. MIR spectroscopy presented in \citet{2000ApJ...535..626I} found an absorption feature at 9.7$\mu$m, consistent with absorption by dust near the central engine. The paper is laid out as follows: in $\S$ 2, we present MIR spectroscopic observations and data reduction of Cyg A made at the 8.2 m {\it Subaru} telescope, and archival data from other telescopes that will allow discussion of the broad band spectral properties of Cyg A. In $\S$ 3, we discuss the spectral features of our spectra and compare with data from the {\it Spitzer} and {\it Keck} telescopes. In $\S$ 4, we present and compare modeling of the MIR spectra, discuss the implications for previous work, and discuss the multi-wavelength SED. In $\S$ 5, we discuss the overall significance of these observations and models. | The high angular resolution MIR spectra are dominated by the 10 $\mu$m silicate absorption feature. The total and central spectra reveal a deeper silicate absorption than previously seen in the {\it Spitzer} spectrum. The detection of the deep silicate feature agrees with previous MIR observations, as well as observations in other wavelength bands which detect the presence of absorbing material in the central kiloparsec of Cyg A. Since the previous spectral observation has a slit perpendicular to ours and yet has similar absorption features, the general distribution of dust around the core of Cyg A appears to peak over the nucleus. Our non-detection of any PAH features agrees with all previous spectra of Cyg A and suggests the powerful central engine of Cyg A destroys the PAH or there is not a significant star formation rate in the central kiloparsec. Our modeling of the {\it Subaru} central spectrum suggest absorption by foreground dust along the line of sight to the torus plays a significant, if not dominant, role in producing the deep silicate feature. The CLUMPY models cannot reproduce a deep silicate absorption feature, and the simple model of a cold dust screen agrees with previous observations of a thick dust lane in Cyg A. Therefore, it is difficult to quantify the contribution of the torus, or any of its properties, from the data. The temperature of the screened blackbody suggests the presence of MIR emission behind the screen. Based in part on data collected at {\it Subaru} Telescope, which is operated by the National Astronomical Observatory of Japan. M.J.M. and E.S.P acknowledge support from NSF grant AST-0904890. C. P. acknowledges support from NFS grant AST-0904421. M. E. acknowledges support from NFS grant AST-0904316. We are pleased to acknowledge the helpful discussions with members of the Los Piratas, especially Rachel Mason and Enrique Lopez Rodriguez for intensive help with data reduction. | 14 | 4 | 1404.0335 |
1404 | 1404.7515_arXiv.txt | {The evolution and spectral properties of stars on the asymptotic giant branch (AGB) are significantly affected by mass loss through dusty stellar winds. Dynamic atmosphere and wind models are an essential tool for studying these evolved stars, both individually and as members of stellar populations, to understand their contribution to the integrated light and chemical evolution of galaxies. } {This paper is part of a series with the purpose of testing state-of-the-art atmosphere and wind models of C-type AGB stars against observations, and making them available to the community for use in various theoretical and observational studies. } {We have computed low-resolution spectra and photometry (in the wavelength range 0.35 -- 25 $\mu$m) for a grid of 540 dynamic models with stellar parameters typical of solar-metallicity C-rich AGB stars and with a range of pulsation amplitudes. The models cover the dynamic atmosphere and dusty outflow (if present), assuming spherical symmetry, and taking opacities of gas-phase species and dust grains consistently into account. To characterize the time-dependent dynamic and photometric behaviour of the models in a concise way we defined a number of classes for models with and without winds. } {Comparisons with observed data in general show a quite satisfactory agreement for example regarding mass-loss rates vs. \mbox{($J$\,--\,$K$)} colours or $K$ magnitudes vs. \mbox{($J$\,--\,$K$)} colours. Some exceptions from the good overall agreement, however, are found and attributed to the range of input parameters (e.g. relatively high carbon excesses) or intrinsic model assumptions (e.g. small particle limit for grain opacities). } {While current results indicate that some changes in model assumptions and parameter ranges should be made in the future to bring certain synthetic observables into better agreement with observations, it seems unlikely that these pending improvements will significantly affect the mass-loss rates of the models. } | \label{s:intro} High luminosities, low effective temperatures and highly dynamic atmospheres are defining properties of AGB stars that make them popular targets for observations ranging from photometric monitoring to high-resolution spectroscopy and, more recently, NIR and MIR interferometry. They represent a crucial stage in the life of low- and intermediate-mass stars where mass loss is a decisive factor for the evolution, and newly produced chemical elements, most notably carbon, are fed into the surrounding ISM in the form of gas and dust by stellar winds. The production of new elements and their transport to the surface are the result of complex interlinked processes involving thermal pulses, hot bottom burning, and the third dredge-up \citep[see, e.g.][]{Herw05, Kara11, StLa11, Marig13}. In addition to being interesting objects in their own right, AGB stars are prominent members of stellar populations and important tools for extragalactic studies. All of these facts are good reasons for developing a better physical understanding of these stars, and a crucial part of that effort is the construction of detailed, realistic models of their atmospheres and winds. This article is the fourth in a series dedicated to an in-depth analysis of synthetic observables, in particular, low-resolution spectra and photometry, resulting from state-of-the-art model atmospheres for carbon-rich giant stars. The first paper \citep[][Paper~I]{AGNML09} presented a large grid of classical hydrostatic model atmospheres for C-type AGB stars, demonstrating that such models are applicable to the interpretation of objects with low pulsation amplitudes and very low mass-loss rates. For effective temperatures above 2800\,K these models agree very well with observations. For cooler, more dynamic stars, however, the formation of dusty circumstellar envelopes severely affects the spectral energy distribution. The time-dependent effects of shock waves and dust formation on observable properties were discussed in detail in the second paper of this series \citep[][Paper~II]{NoAHL11}, based on one dynamic model with an intermediate mass-loss rate. The third paper \citep[][Paper~III]{NoAHE13} extended the study of dynamic models to a small sample with stellar parameters representing different stages of AGB evolution, resulting in a sequence of increasing mass loss and correspondingly stronger circumstellar reprocessing of the photospheric fluxes. A detailed comparison with photometric observations of C-type AGB stars showed that our dynamic atmosphere and wind models reproduce a wide range of observed properties. References to other work based on hydrostatic and dynamic models can be found in Papers~I and II, respectively. In this paper we present results for 540 dynamic models, to a large extent based on the grid by \cite{MatWH10}, to provide a consistent set of wind properties (mass-loss rates, outflow velocities, dust-to-gas ratios) and synthetic observables (low-resolution spectra, photometry) to the community. We discuss the influence of fundamental stellar parameters, pulsation properties, and certain model assumptions on the results, and compare them with various observations of carbon stars. This unique data set and accompanying information is made available in electronic form, providing input for theoretical studies on the evolution of individual stars and stellar populations, as well as for interpretations of observations. Section \ref{s:modelling} gives a short summary of modelling methods, described in more detail in earlier papers, and an overview of the physical parameters defining the model grid. In Sect.~\ref{s:repcases} we define classes characterizing the time-dependent dynamic and photometric behaviour of the individual models in a compact form (with representative examples discussed in Appendix~\ref{a:repremods}). In Sect.~\ref{s:properties} we give an overview of the dynamic and photometric properties of the whole model grid, as well as a comparison with observations. Section~\ref{s:summary} presents a brief summary and our main conclusions. Appendices~\ref{a:overviewdata} and \ref{a:download} contain a short description of the material available at the CDS. | \label{s:summary} The results presented in this paper are based on state-of-the-art atmosphere and wind models of carbon stars, spanning a considerable range in fundamental stellar parameters (given in Table~\ref{t:gridparam}) with three values of carbon excess, three values of piston velocity amplitude, and two values of the parameter $f_{\rm L}$ defining the luminosity amplitude. For these 540 models, we performed {\it a posteriori} radiative transfer calculations for representative snapshots of structures, covering wavelengths from 0.35 to 25 $\mu$m. We considered opacities from gas (atomic and molecular lines, continuous sources) as well as from amorphous carbon dust grains (when present). The results are available for downloading and include dynamic properties (mass-loss rates, outflow velocities, dust condensation degrees) as well as spectra and photometric quantities. The data, especially the mass-loss rates, can be used in stellar evolution modeling, or for population synthesis where photometric properties are needed. Another possibility is to study individual carbon stars using these data. When {\em comparing our results with observed quantities} for individual or samples of carbon stars, several aspects must be kept in mind. First we have the intrinsic limitations and assumptions in the models, such as neglecting 3D effects by assuming spherical symmetry, and using the small-particle limit for grain opacities. Second, certain parameter choices for the current modelling, such as the velocity and luminosity amplitudes, are proxies for a more realistic treatment of the pulsational behaviour of real AGB stars. Third, our grid contains models regardless of how common a particular parameter combination would be in a real stellar population, characterized by an initial mass function, age, and metallicity distribution, and modified by stellar evolution. Despite these caveats, we see, in general, a rather satisfactory agreement of typical models with representative stars, regarding both dynamic and photometric properties, such as mass-loss rates vs. \mbox{($J$\,--\,$K$)} colours, $K$ magnitudes vs. \mbox{($J$\,--\,$K$)} colours, and the colour-colour diagram of \mbox{($J$\,--\,$H$)} vs. \mbox{($H$\,--\,$K$)} (the latter with the exception of models with small or no mass-loss; probably an effect of lacking low (C--O) values in the current grid). This is in line with our earlier detailed comparisons of selected models with individual carbon stars \citep[see Papers II and III;][]{GaHJH04, NowHA10, SAHNPVH10}. Two types of systematic exceptions from the good overall agreement between models and observations should, however, be mentioned here: in various diagrams we note (a) groups of models with no observed counterparts, and (b) ranges of observed quantities not fully covered by models. Both phenomena can, in principle, be caused by the chosen range of parameters of the model grid, or by effects of intrinsic physical assumptions in the models. In particular, we find the following notable discrepancies: (i) our grid does not fully cover the extreme red end of the \mbox{($J$\,--\,$K$)} range in the \cite{WhFMG06} sample, possibly because the highest mass-loss rates in the models are smaller than the highest values derived for that group of stars. In this context, we note that the value of the dust-to-gas ratio commonly assumed to derive mass-loss rates from observations (i.e. 0.005) is significantly higher than the typical values found in our models (cf. Fig.~\ref{fig3}), and that the theoretical values show a large spread. (ii) Models combining low mass and low luminosity with high enough carbon excess and pulsation amplitude to initiate a wind have no counterparts among the observed stars in the diagrams showing mass loss vs. period or $K$ magnitude vs. \mbox{($J$\,--\,$K$)}. In addition, the models with the highest carbon excess used in this grid have a tendency to produce too high wind velocities compared with the bulk of observed stars. A major determining factor for the outflow velocity is the carbon excess, and thus the amount of amC grains formed. This leads us to conclude that stars with such high carbon excess are rare. (iii) In contrast to models at the high end of the assumed (C--O) range, models that do form winds with the lowest (C--O) values tend to show too low outflow velocities compared with their comparatively high mass-loss rates. Based on an exploratory study by \cite{MatH11}, we expect that using size-dependent grain opacities (instead of the small particle limit) will increase the wind speeds of this model group without affecting the mass-loss rates significantly, bringing them into agreement with the bulk of the observed stars in the mass-loss rate vs. wind velocity diagram. As a consequence of the present results, we plan to improve the modelling regarding both intrinsic model assumptions and parameter ranges. In particular, we will provide a full grid based on size-dependent grain opacities, which may both solve current problems with wind velocities, and probably improve synthetic $V$ magnitudes, for example, for which the effects of dust are crucial. Then also smaller carbon excesses will be included in the parameter range of the grid, which is expected to bring the location of models with low or no mass loss in the \mbox{($J$\,--\,$H$)} vs. \mbox{($H$\,--\,$K$)} diagram into better agreement with observations. Conclusions relevant for the {\em application of mass-loss rates in stellar evolution models} can be summarized as follows: the dynamic effects resulting from increasing the luminosity amplitude relative to the grid by \cite{MatWH10} by a factor of 2 for otherwise identical parameters (i.e. similar input of kinetic energy by pulsation) are small, both regarding the mass-loss rates and the outflow velocities. Moreover, as demonstrated in the exploratory study by \cite{MatH11}, using size-dependent grain opacities instead of the small-particle limit probably does not affect the mass-loss rates considerably (despite a noticeable effect on outflow speeds for slow winds). Some exceptions to these conclusions may apply to marginal winds (low velocity and low mass-loss rates), but these are not expected to have a significant effect on evolution models. Therefore we conclude that the mass-loss rates given by \cite{MatWH10} can be used for stellar evolution modelling, even in the light of current and pending improvements affecting synthetic spectra, photometry, and other observables derived from the models. It should also be mentioned here that the mass-loss formula given by \cite{Wacht02} (based on an older generation of dynamical wind models) gives systematically higher values of mass-loss rates for the same combinations of stellar parameters. In the light of the rather good agreement of our model grid with observations, mass-loss rates resulting from the \cite{Wacht02} formula probably need to be revised downwards. | 14 | 4 | 1404.7515 |
1404 | 1404.4450_arXiv.txt | Accreting, nuclear-burning white dwarfs have been deemed to be candidate progenitors of type Ia supernovae, and to account for supersoft X-ray sources, novae, etc. depending on their accretion rates. We have carried out a binary population synthesis study of their populations using two algorithms. In the first, we use the binary population synthesis code \textsf{BSE} as a baseline for the ``rapid'' approach commonly used in such studies. In the second, we employ a ``hybrid'' approach, in which we use \textsf{BSE} to generate a population of white dwarfs (WD) with non-degenerate companions on the verge of filling their Roche lobes. We then follow their mass transfer phase using the detailed stellar evolution code \textsf{MESA}. We investigate the evolution of the number of rapidly accreting white dwarfs (RAWDs) and stably nuclear-burning white dwarfs (SNBWDs), and estimate the type Ia supernovae (SNe Ia) rate produced by ``single-degenerate'' systems (SD). We find significant differences between the two algorithms in the predicted numbers of SNBWDs at early times, and also in the delay time distribution (DTD) of SD SNe Ia. Such differences in the treatment of mass transfer may partially account for differences in the SNe Ia rate and DTD found by different groups. Adopting 100\% efficiency for helium burning, the rate of SNe Ia produced by the SD-channel in a Milky-way-like galaxy in our calculations is $2.0\times10^{-4}\rm{yr}^{-1}$, more than an order of magnitude below the observationally inferred value. In agreement with previous studies, our calculated SD DTD is inconsistent with observations. | \label{sec:intro} Type Ia supernovae (SNe~Ia) have been used with great success as standardizable candles, allowing for the measurement of cosmological parameters (\citealt{rfcc+98,pagk+99}). SNe Ia are also of great importance for galactic chemical evolution \citep[e.g.][]{mg86}. It is widely accepted that they are thermonuclear explosions of carbon-oxygen (CO) white dwarfs (WDs). The compact, degenerate structure of the exploding stars in \sne\ was recently confirmed by early-time multiwavelength observations of SN2011fe \citep{nsct+11,bksb+12}. However, the nature of \sne\ progenitors is still unclear (see \citealt{hkrr13} for a recent review). The models for the progenitors of SN~Ia fall into two categories: the single degenerate (SD) model \citep{wi73} and the double degenerate (DD) model \citep{ty81,it84,webb84}. In the standard SD-model a WD accretes matter from a non-degenerate companion, which may be a main-sequence, subgiant, or red giant star. In order to grow, a WD must accumulate mass via nuclear-burning of hydrogen into helium, and helium into carbon and oxygen. When the WD mass reaches \mch, the WD explodes as an SN~Ia. However, theoretical and observational challenges persist for both scenarios. The fundamental difficulty for the SD-model is the narrow range of accretion rates ($\sim$few$\times10^{-7}$\,\myr) for which steady nuclear-burning and efficient accumulation of mass by the WD is possible \citep{pz78}. This requires specific combinations of donor and accretor masses, restricting the typical delay time between formation of a binary and a \sna\ by $\sim$1 Gyr, and similarly the peak production of \sne\ in this channel within a similar delay time. Another problem is the treatment of the excess matter which cannot be processed through steady nuclear-burning. This is typically \textit{assumed} either to form an extended envelope around the WD, leading to the formation of a common envelope, or to be lost from the system in the form of an optically thick wind. Therefore, the viability of the SD-scenario depends critically on the treatment of mass transfer and resulting accretion rate, which defines whether the WD may, presumably grow in mass. White dwarfs with different accretion rates are associated with different sources and phenomena, e.g. supersoft X-ray sources (SSSs) and novae. Comparing observations with the number of SSSs and the nova rate predicted by population synthesis models can be used to verify calculations, and also to constrain the SD-channel. Because of the relatively high mass transfer rates needed to sustain steady nuclear burning, these sources are almost always associated with mass transfer on the donor's thermal timescale (thermal timescale mass transfer, TTMT). In binary population synthesis codes, TTMT is typically accounted for using a simple analytic treatment. However, such analysis typically assumes implicitly that the donor star remains in {\it thermal equilibrium}, with the entire star (or at least its entire envelope) responding at once, despite mass transfer being driven by the {\it thermal disequilibrium} of the donor \citep[e.g][]{yltk95, rbf09, btn13}. This is particularly important in treating mass loss from red giants -- detailed calculations reveal that the rapid expansion of the donor envelope in response to mass transfer, expected in the simplified treatment of adiabatic models \citep{hw87}, does not necessarily occur \citep{wi11}. This is critical in determining the circumstances under which a binary will undergo a common envelope (CE) phase. In those cases where the binary will undergo a CE regardless, it is also possible that some mass may be accreted prior to this phase, and any accreting WD may appear briefly as an SSS. This is unaccounted for in the traditional treatment of mass transfer in population synthesis. In this paper (Paper I), we investigate in detail mass transfer in the semidetached systems with nuclear-burning WD (NBWD) accretors and main-sequence, Hertzsprung gap and red-giant donors. We pay special attention to the systems in which WDs burn hydrogen steadily (SNBWDs) and to the systems with accretion rates exceeding the upper limit for steady burning, but too low for the formation of a common envelope, \citep[``rapidly accreting white dwarfs'' (RAWDs),][]{lv13} \footnote{It was shown by \citet{1971AcA....21..417P} that putting a $\simeq 10^{-3}$\,\ms\ hydrogen-helium envelope atop a hot ($\log T_e=5.0$) carbon-oxygen WD transforms it into a red giant ($\log T_e=3.6$); this may be avoided, if excess of the matter is removed by postulated optically-thick stellar wind \citep{hkn99}.}. For this, we produce a grid of $\sim3\times10^{4}$ evolutionary sequences of close binary models with different initial combinations of WD accretors and nondegenerate donors, and with differing orbital periods at the onset of Roche lobe overflow, calculated by the detailed stellar evolutionary code \textsf{MESA} \citep{pbdh+11,pcab+13}. Our models are compared with the ones obtained using analytic descriptions of mass-transfer. In order to relate our work to observations, we compare the predicted evolution of the numbers of SNBWD, RAWD, and the rates of \sne\ given two star formation histories: an instantaneous burst of star formation, and a constant star formation rate for 10\,Gyr, approximating early and late type galaxies respectively. In a subsequent paper (hereafter Paper II), we will incorporate spectral models for nuclear-burning white dwarfs. This will allow us to more meaningfully test the predictions of our model. We describe the method of calculations in \S\ref{sec:method}, highlight the effect of varying treatments of TTMT in $\S$\ref{TTMT}, follow with a discussion of how some observables vary with changing MT treatment in \S\ref{sec:res}, in particular the predicted populations of RAWDs, SNBWDs, and \sne. Summary and conclusions are presented in \S\ref{sec:sum}. | \label{sec:sum} In this work, we combined the population synthesis code \textsf{BSE} with the detailed stellar evolutionary code \textsf{MESA} for the first time, in order to study the population of accreting WDs. We also compared the output from this with the results obtained applying a ``rapid'' algorithm, using \textsf{BSE} alone. With these two BPS algorithms, we investigated the evolution of the number of rapidly accreting white dwarfs, stable nuclear burning white dwarfs and the SNe Ia rate in elliptical and spiral-like galaxies. In addition to confirming that the SD channel is subdominant in producing the overall SN Ia rate, we also evaluated the effect of implementing differing treatments of mass transfer for the results of BPS calculations. Comparing the two versions of our binary population synthesis calculations, we found that the mass transfer prescription in BPS is especially important for calculating the number and total luminosity of nuclear-burning white dwarfs in elliptical galaxies at early and late epochs (from initial starburst). We argue that this also partially explains the differences in SNe~Ia rates and DTD obtained by different binary population synthesis groups. We found that RAWDs appear earlier in our \textsf{BSE+MESA} model compared with \textsf{BSE}-only, due to the accreting WDs with massive donors found in the former. We find that there is a factor of $\approx$ 3 difference between the results of our calculations using a fitting formula for the binding parameter $\lambda$ \citep{lvk11} and a constant $\alpha_{\rm ce}=0.25$, and our calculations using a constant $\alpha_{\rm ce} \times \lambda=0.25$. In our \textsf{BSE+MESA} model, we found that the number of RAWDs at 10 Gyr is $160 - 180$ for an elliptical galaxy of $10^{11}$\,\ms\ and $2250 - 2500$ for a spiral-like galaxy of the same mass. This result is in stark contrast with zero RAWDs predicted in our calculation for a model elliptical galaxy using \textsf{BSE} alone. We find that the number of SNBWD is $750 - 1900$ in our model elliptical-like galaxy and $4550 - 6550$ in our model spiral-like galaxy (both with $M=10^{11}$\,\ms\,at 10 Gyr). The predicted SD SNe Ia rate for a Milky-Way-like galaxy is found to be $\simeq2.0\times 10^{-4} \mathrm{yr}^{-1}$, more than an order of magnitude lower than the observationally inferred total Galactic SNe~Ia rate. Our DTD for the SD-model is inconsistent with the observed DTD $\propto t^{-1}$. If we assume that RAWDs do not exist, but rather that a common envelope is formed if the accretion rate onto a WD is larger that the upper limit for stable hydrogen burning, then the rate of \sne\ produced by the SD-channel becomes negligible. To conclude our discussion of the significance of the SD-channel for the production of \sne, we note the following. Since WDs in the steady burning phase can effectively accumulate mass, it is widely suggested that SSSs be identified with the progenitors of SNe Ia and, therefore, observations of SSSs may be useful for constraining the SD model. \citet{gb10} estimated the expected X-ray flux from the progenitors of SNe Ia in the SD-scenario for elliptical galaxies based on the observed supernova rate. They found that the observed X-ray flux from six nearby elliptical galaxies is $30-50$ times smaller than the predicted value, and constrain the contribution of the SD-channel to $<5\%$. In a similar way, \citet{ste10a} found that there are too few SSSs to account for the SNe Ia rate. Even if accreting WDs radiate at significantly lower temperatures (T$\approx 10^5K$), due to the inflation of their photospheres, the SD channel may still be limited to providing $< 10\%$ of the \sne\ rate \citep{2013MNRAS.432.1640W,jwgs+14}. \citet{yltt+96}, \citet{ste10a}, and \citet{y10} considered the possibility that some SSSs may reside in wind-accreting systems which later produce double-degenerate systems (symbiotic stars), and estimated their number in the Galaxy as $\sim 10^3$, while the estimate by \citet{nnvt14} is even lower: $\sim 10^2$. Though several $10^3$\ SNBWDs in the model is apparently a large number, it is still about 2 orders of magnitude lower than that necessary to be consistent with the observationally inferred Galactic \sne\ rate. \citet{2006MNRAS.372.1389L} and \citet{nnvt14} estimated that in wind-fed systems WD typically accrete no more than $\approx$0.1\ms. | 14 | 4 | 1404.4450 |
1404 | 1404.6987_arXiv.txt | To estimate the peculiar velocity of the Sun with respect to the Local Standard of Rest (LSR), we used young objects in the Solar neighborhood with distance measurement errors within 10\%--15\%. These objects were the nearest Hipparcos stars of spectral classes O--B2.5, masers with trigonometric parallaxes measured by means of VLBI, and two samples of the youngest and middle-aged Cepheids. The most significant component of motion of all these stars is induced by the spiral density wave. As a result of using all these samples and taking into account the differential Galactic rotation, as well as the influence of the spiral density wave, we obtained the following components of the vector of the peculiar velocity of the Sun with respect to the LSR: $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}= (6.0,10.6,6.5)\pm(0.5,0.8,0.3)\;\mathrm{km\,s}^{-1}$. We have found that components of the Solar velocity are quite insensitive to errors of the distance $R_0$ in a broad range of its values, from $R_0=7.5$~kpc to $R_0=8.5$~kpc, that affect the Galactic rotation curve parameters. In the same time, the Solar velocity components $(U_\odot)_\mathrm{LSR}$ and $(V_\odot)_\mathrm{LSR}$ are very sensitive to the Solar radial phase $\chi_\odot$ in the spiral density wave. | \label{method} Assuming that the angular rotation velocity of the Galaxy ($\Omega$) depends only on the distance $R$ from the axis of rotation, $\Omega=\Omega(R)$, the apparent velocity ${\bf V}(r)$ of a star at heliocentric radius ${\bf r}$ can be described in vectorial notation by the following relation~(Eq.~2.86 in~\cite{ogor65}) \begin{equation} {\bf V}(r)=-{\bf V}_\odot + {\bf V_\theta}(R)-{\bf V_\theta}(R_0) + {\bf V'}, \label{Bottlinger-01}\end{equation} where ${\bf V}_\odot(U_\odot,V_\odot,W_\odot)$ is the mean stellar sample velocity due to the peculiar Solar motion with respect to the LSR (hence its negative sign), the velocity $U$ is directed towards the Galactic center, $V$ is in the direction of Galactic rotation, $W$ is directed to the north Galactic pole; $R_0$ is the Galactocentric distance of the Sun; $R$ is the distance of an object from the Galactic rotation axis; ${\bf V_\theta}(R)$ is the circular velocity of the star with respect to the center of the Galaxy, ${\bf V_\theta}(R_0)$ is the circular velocity of the Sun, while ${\bf V'}$ are residual stellar velocities. It is necessary to note that Eq.~(\ref{Bottlinger-01}) is widely used by different authors for the Galaxy kinematics analysis (for example Eq.~(6) in~\cite{Mendez2000} or Eq.~(4) in~\cite{Vallenari06}). From the above relation~(\ref{Bottlinger-01}), one can write down three equations in components ($V_r,V_l,V_b$), the so-called \textit{Bottlinger's equations}~(Eq.~6.27 in \cite{Trumpler53}): \begin{equation} \begin{array}{lllll} V_r= (\Omega-\Omega_0)R_0\sin l \cos b,\\ V_l= (\Omega-\Omega_0)R_0\cos l-\Omega r\cos b,\\ V_b=-(\Omega-\Omega_0)R_0\sin l \sin b. \label{Bottl-0234} \end{array} \end{equation} These are exact formulas, and the signs of $\Omega$ follow Galactic rotation. The first of these equations was initially deduced by~\cite{Bottl1931}, while the second one, even earlier, by~\cite{Pilowski1931}, and the one for $V_b,$~--- by~\cite{ogor48}. After expanding $\Omega$ into Taylor series against the small parameter $R-R_0$, then expanding the difference $R-R_0$, where the distance $R$ is \begin{equation} R^2=r^2\cos^2 b-2R_0 r\cos b\cos l+R^2_0, \label{RR} \end{equation} and then substituting the result into Eq.~(\ref{Bottl-0234}), one gets the equations of the Oort--Lindblad model (Eq.~6.34 in \cite{Trumpler53}). Our approach departs from the above in that the distances $r$ are known quite well. In this case, there is no need to expand $R-R_0$ into series, since the distance $R$ is calculated from Eq.~(\ref{RR}). Furthermore, our approach implies an extra assumption that the observed stellar velocities include perturbations due to the spiral density wave ${\bf V}_{sp}(V_R,\Delta V_\theta)$, with a linear dependence on both ${\bf V}_{sp}$ and ${\bf V}_\odot$. This allows us to write \begin{equation} -{\bf V}_\odot=-{{\bf V}_\odot}_\mathrm{LSR}+{\bf V}_{sp}. \label{LSR}\end{equation} Perturbations from the spiral density wave have a direct influence on the peculiar Solar velocity ${{\bf V}_\odot}_\mathrm{LSR}$, which appears to be first pointed out by~\cite{Creze1973}~--- see the coefficients $a_1$ and $a_2$ in Eq.~(22) of their paper. Then the relation~(\ref{Bottlinger-01}) takes the following form: \begin{equation} {\bf V}(r)=-{{\bf V}_\odot}_\mathrm{LSR}+{\bf V}_{sp} + {\bf V_\theta}(R)-{\bf V_\theta}(R_0) + {\bf V'}, \label{Bottlinger-02}\end{equation} which, considering the expansion of the angular velocity of Galactic rotation $\Omega$ into series up to the second order of $r/R_0$ reads \begin{equation} \begin{array}{lll} V_r=-U_\odot\cos b\cos l-V_\odot\cos b\sin l-W_\odot\sin b\\ +R_0(R-R_0)\sin l\cos b \Omega^\prime_0\\ +0.5R_0 (R-R_0)^2 \sin l\cos b \Omega^{\prime\prime}_0\\ +\Delta V_{\theta}\sin(l+\theta)\cos b-V_R \cos(l+\theta)\cos b, \label{EQ-1} \end{array} \end{equation} \begin{equation} \begin{array}{lll} V_l= U_\odot\sin l-V_\odot\cos l\\ +(R-R_0)(R_0\cos l-r\cos b) \Omega^\prime_0\\ +(R-R_0)^2 (R_0\cos l - r\cos b)0.5\Omega^{\prime\prime}_0- r \Omega_0 \cos b\\ +\Delta V_{\theta} \cos(l+\theta)+V_R\sin(l+\theta), \label{EQ-2} \end{array} \end{equation} \begin{equation} \begin{array}{lll} V_b=U_\odot\cos l \sin b + V_\odot\sin l \sin b-W_\odot\cos b\\ -R_0(R-R_0)\sin l\sin b\Omega^\prime_0\\ -0.5R_0(R-R_0)^2\sin l\sin b\Omega^{\prime\prime}_0\\ -\Delta V_{\theta} \sin (l+\theta)\sin b+V_R \cos(l+\theta)\sin b, \label{EQ-3} \end{array} \end{equation} where the following designations are used: $V_r$ is the line-of-sight velocity, $V_l=4.74 r \mu_l\cos b$ and $V_b=4.74 r \mu_b$ are the proper motion velocity components in the $l$ and $b$ directions, respectively, with the factor~4.74 being the quotient of the number of kilometers in an astronomical unit and the number of seconds in a tropical year; the star's proper motion components $\mu_l\cos b$ and $\mu_b$ are in mas~yr$^{-1}$, and the line-of-sight velocity $V_r$ is in km~s$^{-1}$; $\Omega_0$ is the angular velocity of rotation at the distance $R_0$; parameters $\Omega^\prime_0$ and $\Omega^{\prime\prime}_0$ are the first and second derivatives of the angular velocity, respectively. To account for the influence of the spiral density wave, we used the simplest kinematic model based on the linear density wave theory by~\cite{LinShu64}, where the potential perturbation is in the form of a travelling wave. Then, \begin{equation} V_R=f_R \cos \chi, \label{VR-VR} \end{equation} \begin{equation} \Delta V_{\theta}=f_\theta \sin \chi, \label{VR-Vtheta} \end{equation} where $f_R$ and $f_\theta$ are the amplitudes of the radial (directed toward the Galactic center in the arm) and azimuthal (directed along the Galactic rotation) velocity perturbations; $i$ is the spiral pitch angle ($i<0$ for winding spirals); $m$ is the number of arms (we take $m=2$ in this paper); $\theta$ is the star's position angle measured in the direction of Galactic rotation: $\tan\theta = y/(R_0-x)$, where $x$ and $y$ are the Galactic heliocentric rectangular coordinates of the object; radial phase of the wave $\chi$ is \begin{equation} \chi=m[\cot (i)\ln (R/R_0)-\theta]+\chi_\odot, \label{chi-creze} \end{equation} where $\chi_\odot$ is the radial phase of the Sun in the spiral density wave; we measure this angle from the center of the Carina--Sagittarius spiral arm ($R\approx7$~kpc). The parameter $\lambda$, which is the distance along the Galactocentric radial direction between adjacent segments of the spiral arms in the Solar neighborhood (the wavelength of the spiral density wave), is calculated from the relation $2\pi R_0/\lambda = m\cot(i).$ \begin{figure*} \includegraphics[width=2.0\columnwidth]{f1.eps} \caption{Radial ($V_R$, dark) and tangential ($\Delta V_\theta$, light) perturbation velocities versus Galactocentric distances $R$. Location of the Sun is indicated by a dotted line.} \label{f1} \end{figure*} \begin{figure} \begin{center} \includegraphics[width=0.85\columnwidth]{f2.eps} \caption{Number of stars versus Galactocentric distances $R$.} \label{f2} \end{center} \end{figure} We take $R_0=8.0\pm0.4$~kpc, according to analysis of the most recent determinations of this quantity in the review by~\cite{FosterR010}. We use the well-known statistical method~\citep{Trumpler53,ogor65} to determine the parameters of the residual velocity (Schwarzschild) ellipsoid. It consists in determining the symmetric tensor of moments or the tensor of residual stellar velocity dispersions. When simultaneously using the stellar line-of-site velocities and proper motions to find the six unknown components of the dispersion tensor, we have six equations for each star. The semiaxes of the residual velocity ellipsoid, denoted by $\sigma_{1,2,3}$, can be determined by analyzing the eigenvalues of the dispersion tensor. In the present paper, we assume that parameters of both the differential Galactic rotation and the spiral density wave are known from observations of distant stars and solving equations of the form~(\ref{EQ-1})--(\ref{EQ-3}). In this case the right-hand parts of the equations contain only components of the Solar peculiar velocity \begin{equation} \begin{array}{lll} V_r-R_0(R-R_0)\sin l\cos b \Omega^\prime_0\\ -0.5R_0 (R-R_0)^2 \sin l\cos b \Omega^{\prime\prime}_0\\ -\Delta V_{\theta}\sin(l+\theta)\cos b-V_R \cos(l+\theta)\cos b\\ =-U_\odot\cos b\cos l-V_\odot\cos b\sin l-W_\odot\sin b, \label{EQ-101} \end{array} \end{equation} \begin{equation} \begin{array}{lll} V_l-(R-R_0)(R_0\cos l-r\cos b) \Omega^\prime_0\\ -(R-R_0)^2 (R_0\cos l - r\cos b)0.5\Omega^{\prime\prime}_0+r \Omega_0 \cos b\\ -\Delta V_{\theta} \cos(l+\theta)-V_R\sin(l+\theta)\\ =U_\odot\sin l-V_\odot\cos l, \label{EQ-102} \end{array} \end{equation} \begin{equation} \begin{array}{lll} V_b+R_0(R-R_0)\sin l\sin b\Omega^\prime_0\\ +0.5R_0(R-R_0)^2\sin l\sin b\Omega^{\prime\prime}_0\\ +\Delta V_{\theta} \sin (l+\theta)\sin b+V_R \cos(l+\theta)\sin b\\ =U_\odot\cos l \sin b + V_\odot\sin l \sin b-W_\odot\cos b.\\ \label{EQ-103} \end{array} \end{equation} The system~(\ref{EQ-101})--(\ref{EQ-103}) can be solved by least-squares adjustment with respect to three unknowns $U_\odot,$ $V_\odot$, and $W_\odot$. Another approach (which we follow) is to calculate components of spatial velocities $U,V,W$ of stars: \begin{equation} \begin{array}{lll} U=V'_r\cos l\cos b-V'_l\sin l-V'_b\cos l\sin b,\\ V=V'_r\sin l\cos b+V'_l\cos l-V'_b\sin l\sin b,\\ W=V'_r\sin b +V'_b\cos b, \label{EQ-UVW} \end{array} \end{equation} where $V'_r,V'_l,V'_b$ are left-hand parts of Eqs.~(\ref{EQ-101})--(\ref{EQ-103}) which are the observed stellar velocities free from Galactic rotation and the spiral density wave. Then ${\overline U=-U_\odot},$ ${\overline V=-V_\odot}$ and ${\overline W=-W_\odot}$. | \label{Results} \subsection{The fixed value of $R_0$}\label{Approach1} Here we describe the results obtained at fixed value of $R_0=8$ kpc, assuming the parameters of differential Galactic rotation and the spiral density wave calculated earlier independently for each stellar sample. In Figure~\ref{f1}, there are radial ($V_R$) and tangential ($\Delta V_\theta$) velocities of perturbations vs Galactocentric distance $R$, induced by the spiral density wave. These velocities are calculated according to the formulas (\ref{VR-VR}), (\ref{VR-Vtheta}), and (\ref{chi-creze}) assuming $\theta=0^\circ$, and the amplitudes of perturbations $f_R$ and $f_\theta$ defined in the data description (Section~\ref{Data}). As it can be seen from this figure, at $R=R_0$, the perturbations achieve about $5$~km~s$^{-1}$ in the radial direction. In the tangential direction, the same value is achieved for two samples: of youngest O--B2 stars and of masers. In the case of young Cepheids, perturbations in the tangential direction are not significant. In the case of middle-aged Cepheids, perturbations in the tangential direction at $R=R_0$ are close to zero. Note that a very small Solar neighborhood ($R\rightarrow R_0$) is crucial to determine the velocity $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}$. In Table~\ref{t1}, the components of the peculiar velocity of the Sun with respect to the LSR $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}$ are given. They were obtained only taking into account the influence of the differential Galactic rotation. Components of this vector, given in Table~\ref{t2}, were calculated considering both the effects of the differential Galactic rotation and of the spiral density wave. In the last three columns of Tables~\ref{t1}--\ref{t2}, the main axes of the ellipsoid of residual velocities $\sigma_1,\sigma_2,\sigma_3$ are given. We should note that, after considering the perturbations from the spiral density wave, the values of residual velocity dispersions $\sigma_1,\sigma_2,\sigma_3$ have slightly different distributions, which leads to a changed orientation of the residual velocity ellipsoid. However, a detailed analysis of this problem is out of the scope of the present study and will be conducted elsewhere. As it is seen from Tables~\ref{t1} and \ref{t2}, considering the effect of the spiral density wave for O--B2.5 stars and for masers leads to a considerable variation of the components $\Delta U_\odot$ and $\Delta V_\odot$ by $\approx$6~km$^{-1}$. In addition, this gives smaller errors of the velocity $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}$, which is especially noticeable for masers. The velocity $(V_\odot)_\mathrm{LSR}$ (Table~\ref{t1}) found from the data on masers differs from $(V_\odot)_\mathrm{LSR}=12.2\;\mathrm{km\,s}^{-1}$\citep{Schonrich10} by $\approx$4~km\,s$^{-1}$, which is in accordance with the result of analysis of masers in the Local arm~\citep{Xu13}. The following average values of the parameters $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}$ found in the present work are, essentially, more accurate than the estimate $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}= (5.5,11.0,8.5)\pm(2.2,1.7,1.2)\;\mathrm{km\,s}^{-1}$ obtained from 28 masers by~\cite{bb10} considering the influence of the spiral density wave. The average value of $(V_\odot)_\mathrm{LSR}$ (Table~\ref{t2}) is in a good agreement with the result by~\cite{Schonrich10}. There is a discrepancy in the $(U_\odot)_\mathrm{LSR}$ component with~\cite{Schonrich10}, and especially with~\cite{francis12}. Note that the revised Str$\ddot{o}$mberg relation applied to the experimental RAVE data gives an absolutely different velocity $(V_\odot)_\mathrm{LSR}\approx3\;\mathrm{km\,s}^{-1}$~\citep{Golubov13}. Using another approach to analysis of RAVE data \cite{Pasetto12} obtained the following velocities: $(U_\odot,V_\odot)_\mathrm{LSR}= (9.87,8.01)\pm(0.37,0.29)\;\mathrm{km\,s}^{-1}$. Thus different methods give different results, and a final agreement on the values of the velocity $(U_\odot,V_\odot,W_\odot)_\mathrm{LSR}$ is not achieved till now. We consider our estimates most reliable as they are based on the youngest stars characterized by a small velocity dispersion and by small Galactic orbit eccentricities as well. In Fig.~\ref{f3}, there is a histogram of Galactic orbital eccentricities for the OB-2.5 star sample. One can see that eccentricities of the stars considered are indeed small. Note that we have excluded escaping stars when making this sample~\citep{bb13OB2}. \begin{figure} \begin{center} \includegraphics[width=0.85\columnwidth]{f3.eps} \caption{Number of stars versus eccentricity $e$.} \label{f3} \end{center} \end{figure} \begin{table*} \begin{center} \caption{Components of the vector of the peculiar velocity of the Sun with respect to the LSR, calculated considering both the differential Galactic rotation and the spiral density wave for three values of $R_0=7.5,8.0,8.5$ kpc. } \label{t3} {\small \begin{tabular}{|l|c|c|c|||c|l|c|c|c|c|c|c|c|} \hline $R_0$ & & $7.5$~kpc & && & $8.0$~kpc & && & $8.5$~kpc & \\\hline Stars & $U_\odot$ & $V_\odot$ & $W_\odot$ && $U_\odot$ & $V_\odot$ & $W_\odot$ && $U_\odot$ & $V_\odot$ & $W_\odot$ \\ & km s$^{-1}$ & km s$^{-1}$ & km s$^{-1}$ && km s$^{-1}$ & km s$^{-1}$ & km s$^{-1}$ && km s$^{-1}$ & km s$^{-1}$ & km s$^{-1}$ \\\hline O--B2.5 & $4.9\pm0.5$ & $ 8.6\pm0.5$ & $7.2\pm0.3$ && $4.8\pm0.5$ & $ 8.6\pm0.5$ & $7.2\pm0.3$ && $4.8\pm0.5$ & $ 8.5\pm0.5$ & $7.2\pm0.3$ \\ masers & $5.5\pm1.5$ & $11.3\pm2.5$ & $6.2\pm1.7$ && $6.0\pm1.6$ & $11.4\pm2.5$ & $6.2\pm1.7$ && $5.4\pm1.5$ & $11.5\pm2.5$ & $6.2\pm1.7$ \\ Ceph., $P\geq9^\mathrm{d}$ & $6.1\pm2.3$ & $10.5\pm2.2$ & $6.1\pm2.5$ && $6.6\pm2.5$ & $10.6\pm2.2$ & $6.1\pm2.5$ && $7.0\pm2.6$ & $10.7\pm2.2$ & $6.1\pm2.5$ \\ $5^\mathrm{d}\leq P<9^\mathrm{d}$ & $7.0\pm2.1$ & $10.8\pm1.9$ & $6.4\pm1.8$ && $6.7\pm2.1$ & $10.8\pm2.0$ & $6.4\pm1.8$ && $6.5\pm2.1$ & $10.9\pm2.0$ & $6.4\pm1.8$ \\\hline \end{tabular}} \end{center} \end{table*} \begin{table*} \begin{center} \caption{Components of the vector of the peculiar velocity of the Sun with respect to the LSR, calculated considering both the differential Galactic rotation and the spiral density wave for the different values of $\chi_\odot.$ } \label{t4} {\small \begin{tabular}{|l|c|c|c|||c|l|c|c|c|c|c|c|c||c|c|} \hline $\chi_\odot$ &\multicolumn{2}{|c|} {$-80^\circ$}&\multicolumn{2}{|c|} {$-110^\circ$}&\multicolumn{2}{|c|} {$-120^\circ$}&\multicolumn{2}{|c|} {$-130^\circ$}&\multicolumn{2}{|c|} {$-160^\circ$} \\\hline & $U_\odot$ & $V_\odot$& $U_\odot$& $V_\odot$& $U_\odot$& $V_\odot$& $U_\odot$& $V_\odot$& $U_\odot$& $V_\odot$ \\ & km s$^{-1}$ & km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$& km s$^{-1}$ \\\hline O--B2.5 & $11.4\pm0.5$ & $7.6\pm0.6$& $6.3\pm0.5$& $ 8.0\pm0.6$& $4.8\pm0.5$& $ 8.6\pm0.5$& $3.5\pm0.5$& $ 9.3\pm0.3$& $0.8\pm0.6$& $12.5\pm0.5$ \\\hline \end{tabular}} \end{center} \end{table*} \subsection{Errors of Galactic Rotation Parameters} Here we describe the results obtained for three particular values of $R_0=7.5,8.0,8.5$~kpc using the corresponding differential Galactic rotation parameters. That is, we now use one and the same Galactic rotation curve to analyze each of the stellar samples. Amplitudes of perturbation velocities of the spiral density wave $f_R$ and $f_\theta$, as well as the values of the Solar phase $\chi_\odot$ in the spiral wave, are chosen as above in Section~\ref{Approach1}. For this purpose, we took a sample of masers (55 masers, $\sigma_\pi/\pi<10\%,$ $r<3.5$~kpc) from~\cite{BobBajk2013Masers} and, taking three fixed values of $R_0$, found the following parameters of the Galactic rotation curve: \begin{equation} \begin{array}{lll} R_0= 7.5~\hbox {kpc,}\\ \Omega_0= 30.0\pm1.1~\hbox {km s$^{-1}$ kpc$^{-1},$}\\ \Omega^\prime_0=-4.61\pm0.21~\hbox {km s$^{-1}$ kpc$^{-2},$}\\ \Omega^{\prime\prime}_0=1.081\pm0.180~\hbox {km s$^{-1}$ kpc$^{-3},$} \label{R0-7.5} \end{array} \end{equation} \begin{equation} \begin{array}{lll} R_0= 8.0~\hbox {kpc,}\\ \Omega_0= 29.9\pm1.1~\hbox {km s$^{-1}$ kpc$^{-1},$}\\ \Omega^\prime_0=-4.27\pm0.20~\hbox {km s$^{-1}$ kpc$^{-2},$}\\ \Omega^{\prime\prime}_0=0.915\pm0.166~\hbox {km s$^{-1}$ kpc$^{-3},$} \label{R0-8} \end{array} \end{equation} \begin{equation} \begin{array}{lll} R_0= 8.5~\hbox {kpc,}\\ \Omega_0= 29.8\pm1.1~\hbox {km s$^{-1}$ kpc$^{-1},$}\\ \Omega^\prime_0=-3.98\pm0.18~\hbox {km s$^{-1}$ kpc$^{-2},$}\\ \Omega^{\prime\prime}_0=0.783\pm0.154~\hbox {km s$^{-1}$ kpc$^{-3}.$} \label{R0-8.5} \end{array} \end{equation} Parameters~(\ref{R0-8}) are the same as those used previously for the maser sample in Section~\ref{Approach1}. The results are summarized in Table~\ref{t3}, where one can see that there is no considerable departure from the previous results (Table~\ref{t2}). The only noticeable difference is for the sample of young Cepheids, $\Delta V_\odot\approx2$~km s$^{-1}$, which is due to the difference in Galactic rotation rate $\Delta \Omega_0\approx2$~km s$^{-1}$. As early as in the paper of~\cite{FeastW97} it was already noted that the youngest Cepheids, for some unknown reason, rotate slightly slower than the older ones. In this sense, we consider the approach taken in the previous paragraph as more adequate to the goal of the present study: it is better to apply individual rotation curves to each stellar sample. \subsection{Errors of the Spiral Wave Parameters} Here we describe our results for several model values of the Solar phase $\chi_\odot$ in the spiral density wave for a sample of O--B2.5 stars (161 stars, $r<0.7$~kpc). We used the Galactic rotation curve parameters~(\ref{R0-8}). The results are reflected in Table~\ref{t4} whence one can see that the Solar velocity components $U_\odot$ and $V_\odot$ are very sensitive to the above parameter ($W_\odot$ velocities are not shown in the Table as they are practically not affected by the density wave). It is easy to understand these results by analyzing the corresponding panel of Fig.~\ref{f1} and Table~\ref{t1}. For instance, for $\chi_\odot=-160^\circ$, the radial perturbation curve ($V_R$) is near its maximum, so the influence to the $U_\odot$ component is most prominent. On the contrary, the tangential perturbation curve ($\Delta V_\theta$) is about zero, so there is no effect on the $V_\odot$ component. We must note that, in our previous paper~\citep{bb13OB2}, the uncertainty of $R_0$ was not taken into account when determining the Solar phase in the spiral density wave $\chi_\odot=-120\pm4^\circ$. We have redone Monte Carlo simulation and obtained the following results: \begin{enumerate} \item If we consider only the error $\sigma_{R_0}=0.4$~kpc, its effect on the uncertainty of the Solar phase in the spiral density wave becomes very small: $\sigma_{\chi_\odot}=0.2^\circ$. The explanation for this is that when you change $R_0,$ the length of a wave stretches like a rubber band, but the phase of the Sun in the spiral wave practically does not change. \item If we consider the errors of all observed parameters of stars~-- parallaxes, proper motions, line-of-site velocities~-- along with the uncertainty $\sigma_{R_0}$, then the Solar phase in the spiral density wave becomes $\chi_\odot=-120\pm6^\circ$. \end{enumerate} Based on the data from Table~\ref{t4}, we may conclude that, in the range of phase values from $-110^\circ$ to $-130^\circ$ (which is even above the $1\sigma$ level), the Solar velocities in question, found from O--B2.5 stars, are in the $U_\odot=6-4$~km~s$^{-1}$ and $V_\odot=8-9$~km~s$^{-1}$ range. | 14 | 4 | 1404.6987 |
1404 | 1404.7273_arXiv.txt | We report multi-epoch VLBI H$_2$O maser observations towards the compact cluster of YSOs close to the Herbig Be star LkH$\alpha$~234. This cluster includes LkH$\alpha$~234 and at least nine more YSOs that are formed within projected distances of $\sim$10~arcsec ($\sim$9,000~au). We detect H$_2$O maser emission towards four of these YSOs. In particular, our VLBI observations (including proper motion measurements) reveal a remarkable very compact ($\sim$0.2~arcsec = $\sim$180~au), bipolar H$_2$O maser outflow emerging from the embedded YSO VLA~2. We estimate a kinematic age of $\sim$40~yr for this bipolar outflow, with expanding velocities of $\sim$20~km~s$^{-1}$ and momentum rate $\dot M_w V_w$ $\simeq$ $10^{-4}-10^{-3}$ M$_{\odot}$~yr$^{-1}$~km~s$^{-1}$$\times (\Omega$/$4\pi)$, powered by a YSO of a few solar masses. We propose that the outflow is produced by recurrent episodic jet ejections associated with the formation of this YSO. Short-lived episodic ejection events have previously been found towards high-mass YSOs. We show now that this behaviour is also present in intermediate-mass YSOs. These short-lived episodic ejections are probably related to episodic increases in the accretion rate, as observed in low-mass YSOs. We predict the presence of an accretion disk associated with VLA~2. If detected, this would represent one of the few known examples of intermediate-mass stars with a disk-YSO-jet system at scales of a few hundred au. | The early stages of evolution of low-mass stars are relatively well characterised by the formation of a system with a central protostar, accreting material from a rotating accretion disk at scales of a few hundreds of astronomical units (au), and simultaneously ejecting a collimated outflow with the presence and crucial role of magnetic fields. Accretion and mass-loss processes, two closely related mechanisms, govern the formation of low-mass stars (e.g., Girart, Rao \& Marrone 2006; McKee \& Ostriker 2007; Machida, Inutsuka \& Matsumoto 2008; Armitage 2011; Williams \& Cieza 2011). In fact, spatio-kinematical studies of jets and Herbig-Haro (HH) objects show that outflows from low-mass young stellar objects (YSOs) are non-steady, but presenting variability in the ejection velocity as well as pulsed events, probably related to recurrent instabilities in the accretion disks (e.g., Zinnecker, McCaughrean \& Rayner 1998; Reipurth \& Bally 2001; Estalella et al. 2012). With respect to the formation of massive stars ($\gtrsim$ 10~M$_{\odot}$), there are several examples showing the presence of massive disk-protostar-jet systems at scales of thousand au (in some cases with magnetic fields oriented parallel to the collimated outflows), indicating that stars at least up to $\sim$20~M$_{\odot}$ form via an accretion disk as low-mass do (e.g., Patel et al. 2005; Jim\'enez-Serra et al. 2007; Torrelles et al. 2007; Davies et al. 2010; Carrasco-Gonz\'alez et al. 2010, 2012a; Vlemmings et al. 2010; Fern\'andez-L\'opez et al. 2011; Maud \& Hoare 2013). Furthermore, variation in the ejection velocity of outflows as well as remarkable short-lived episodic ejection events (few tens of years) have been reported towards massive YSOs through radio continuum (e.g., Mart\'{\i} et al 1995; Curiel et al. 2006) and H$_2$O maser spatio-kinematical observations (e.g., Torrelles et al. 2001, 2003; Surcis et al. 2011a,b; Chibueze et al. 2012; Sanna et al. 2012; Kim et al. 2013; Trinidad et al. 2013). This indicates that outflows in massive YSOs are also non-steady, probably due to instabilities in the accretion processes as in the case of low-mass YSOs. Mainly because most of the observational efforts have been concentrated in low- and high-mass YSOs, little is known on what happens in intermediate-mass protostars, with only a few examples showing clear disk-YSO-jet systems at scales of a few hundred au (e.g., Palau et al. 2011; Carrasco-Gonz\'alez et al. 2012b), but up to now without any observational information on the magnetic field and outflow kinematical behaviour at these small scales. Observations of the first stages of the evolution of intermediate-mass YSOs at small scales are therefore important to fill that gap, and thus to have a more complete vision about the processes of star formation throughout a more continuous range of masses. In this paper, we present multi-epoch Very Long Baseline Interferometry (VLBI) H$_2$O maser observations towards intermediate-mass YSOs to study the spatio-kinematical distribution of their masers with $\sim$0.5~mas resolution. This kind of study has previously been very useful in massive star-forming regions to detect outflow activity through the presence of H$_2$O maser emission, allowing the identification of new centres of star formation, some of them associated, as mentioned before, with short-lived episodic ejection events with kinematic ages of a few tens of years. With the present work, we now extend this study to intermediate-mass star-forming regions. The optically visible Herbig Be star LkH$\alpha$~234, with luminosity $\sim$10$^3$~L$_{\odot}$ and mass $\sim$5~M$_{\odot}$, is one of the brightest sources of the NGC~7129 star-forming region, located at a distance of 0.9~kpc (Herbig 1960; Strom et al. 1972; Bechis et al. 1978; Fuente et al. 2001; Trinidad et al. 2004; Xu et al. 2013). High angular resolution near- and mid-infrared images obtained by Kato et al. (2011) reveal, in addition to LkH$\alpha$~234, a cluster of eight YSO candidates within a distance of $\sim$10 arcsec from LkH$\alpha$~234 (named by Kato et al. as objects B, C, D, E, F, G, NW1, and NW2). Very Large Array (VLA) observations show five radio continuum sources in a region of $\sim$5 arcsec (VLA~1, VLA~2, VLA~3A, VLA~3B, LkH$\alpha$~234; Trinidad et al. 2004), some of them with mid-infrared counterparts: VLA~2 (NW2), VLA~3A+3B (NW1), and LkH$\alpha$~234 (see Fig. 1). NW2 and NW1, are not seen in the J, H, and K-band images, but they are bright in the $L'$, $M'$, and mid-infrared bands, indicating that they are highly embedded YSOs (Kato et al. 2011). H$_2$O maser emission is detected towards this region, mainly associated with VLA~1, VLA~2 (NW2), and VLA~3A+3B (NW1) (Tofani et al. 1995; Umemoto et al. 2002; Trinidad et al. 2004; Marvel 2005; see Fig. 1). Several outflows emerging from this region have been also observed, e.g.: i) a jet observed in [S II], extending to the south-west (p.a. $\sim$ 252$^{\circ}$) up to distances of $\sim$40 arcsec ($\sim$36,000~au), without evidence of a counter-jet (Ray et al. 1990); ii) an infrared jet seen in v=1--0 S(1) H$_2$ emission, extending towards the south-west (p.a. $\sim$ 226$^{\circ}$) up to distances of $\sim$10 arcsec ($\sim$9,000~au), also without any evidence of a counter-jet (Cabrit et al. 1997). The overlapping of these two outflows within a small area of the sky makes it very difficult to clearly distinguish between various candidate driving sources. What we know, however, is that all the different studies exclude that the Herbig Be star LkH$\alpha$~234 is driving any of the outflows of this star-forming region, but that the main outflow activity centres are probably the YSOs VLA~2 (NW2) and VLA~3A+3B (NW1) (Trinidad et al. 2004; Marvel 2005; Kato et al. 2011). Following the nomenclature of Trinidad et al. (2004), in this paper we will refer to the region containing all this cluster of YSOs as the ``star-forming LkH$\alpha$~234 region". \begin{figure*} \centering \includegraphics[width=140mm, clip=true]{FIG1_FINAL.eps} \caption{JHK colour composite image ({\it top left-hand panel}), and 11.8 $\mu$m and 17.6~$\mu$m Keck LWS images ({\it top right-hand panels}) around the Herbig Be star LkH$\alpha$~234. In addition to LkH$\alpha$~234, eight YSO candidates are also indicated (B, C, D, E, F, G, NW1, NW2; images and nomenclature from Kato et al. 2011). {\it Bottom panel}: VLA-3.6~cm continuum contour map of the region showing emission from four sources, VLA~1, VLA~2, VLA~3 (this source, when observed with higher angular resolution, has two components, VLA~3A and 3B, which are separated by $\sim$0.3~arcsec; Trinidad et al. 2004), and LkH$\alpha$~234 (the optical position is indicated by a star). The crosses indicate the position of the H$_2$O masers detected with the VLA (epoch 1999 June 29; observations and map from Trinidad et al. 2004). The mid-infrared sources NW1 and NW2 detected by Kato et al. (2011) coincide in position with the radio continuum sources VLA~3 (A+B) and VLA~2, respectively.} \end{figure*} In Section~2, we present the multi-epoch VLBI H$_2$O maser observations towards the star-forming LkH$\alpha$~234 region and the main results, in particular the discovery of a very young, compact bipolar H$_2$O maser outflow associated with VLA~2 (NW2). The implications of these results are discussed in Section~3, while the main conclusions of this work are presented in Section~4. | We report VLBI H$_2$O maser observations towards the intermediate-mass star-forming LkH$\alpha$~234 region. We have detected H$_2$O maser emission towards four YSOs (VLA~1, VLA~2, VLA~3A, and VLA~3B) within a region of $\sim$4~arcsec ($\sim$3,600~au) in size, indicating outflow activity from these objects. In particular, the spatio-kinematical distribution of the masers reveal a remarkable very young, compact bipolar H$_2$O maser outflow associated with the intermediate-mass YSO VLA~2, which is located at the centre of the outflow. This outflow is formed by two bow-shock-like structures moving in opposite directions. The size of the outflow (0.2 arcsec = 180~au; measured from VLA~2), together with the expanding velocity of the masers with respect to the central source ($\sim$20~km~s$^{-1}$), gives a kinematic age of $\sim$40~yr. We interpret this outflow as driven by a short-lived episodic jet ejection from VLA~2, rather than a steady outflow. Short-lived episodic ejections have been also observed in massive YSOs through VLBI H$_2$O maser observations. With the results presented in this paper we show now that this behaviour may also occur in intermediate-mass YSOs, probably due to episodic increases in the accretion rate as observed and firmly established in low-mass YSOs. In this way, our observations predict the presence of an accretion disk around VLA~2 forming a disk-YSO-jet system similar to what is observed in low- and high-mass YSOs. This might be studied with sensitive radio continuum and (sub)mm dust continuum and spectral line observations at scales of a few tenths of an arcsec. Our results support VLA~2 as the powering source of the extended outflow ($\sim$100~arcsec = 9,000~au) seen in [SII]/CO. It is possible that this extended outflow, with kinematic age $\gtrsim$ 8$\times$10$^3$~yr, might be driven by short-lived episodic ejections from VLA~2 as the very recent one traced by the VLBI H$_2$O maser observations. | 14 | 4 | 1404.7273 |
1404 | 1404.5795_arXiv.txt | Polarimetry using a half-wave plate (HWP) modulator provides the strong tools to avoid a detector $1/f$ noise and instrument-originated spurious polarization systematic effects. While the Pancharatnam achromatic HWP (AHWP) is commonly used for an application that needs a broadband frequency coverage, this technique introduces a frequency-dependent polarization angle rotation. In this paper we propose a new technique to mitigate this effect by introducing a second set of an AHWP. One rotational and one stationary set of AHWPs achieve a broadband coverage of modulation efficiency without the frequency-dependent polarization angle rotation. We conducted measurements by using three layers of sapphire wave plates and demonstrated this technique at millimeter wavelengths between 72 and 162~GHz. We also discuss a potential application in the CMB polarization experiment based on numerical simulations. | Measurements of the cosmic microwave background radiation (CMB) have been playing an important role to establish the $\Lambda$CDM cosmology. While vast amount of information is learned by the Planck satellite using the temperature anisotropy of the CMB~\cite{planck}, there is a community-wide effort to measure the polarization of CMB to test the inflationary paradigm and to probe the evolution of the universe via interactions between the CMB and gravitational potential from the large scale structures~\cite{Kamionkowski}. These paradigms started to be disclosed by the recent results from the CMB polarization experiments, BICEP2, POLARBEAR, and SPTpol~\cite{bicep2,polarbear_BB,sptpol}. Forthcoming CMB polarization experiments require a polarimeter that is free from a detector $1/f$ noise and controls the instrument-originated polarization systematic effects. A polarimeter using a half-wave plate (HWP) modulator provides two attractive features, i) avoiding the detector $1/f$ by modulating and demodulating the signal frequency, and ii) eliminating the detector differencing, which is a source of the instrumentally induced spurious polarization effect, to reconstruct the incident polarization state~\cite{hhz}. MAXIPOL was the first CMB experiment that employed the continuously rotating HWP. Currently a number of CMB experiments, including ABS, EBEX, LiteBIRD, POLARBEAR-1, POLARBEAR-2, QUBIC, SPIDER, SWIPE are pursuing this technology~\cite{maxipol,brad_thesis,abs,ebex,litebird,polarbear1,polarbear2,qubic,spider,swipe}. While the use of HWP becomes a popular polarimetry technique, a polarimetry using a single HWP limits the use of the electromagnetic frequency range, and thus the observing detection bandwidth. The typical available bandwidth with a single HWP is $\delta \nu/\nu \sim 0.3$ in order to maintain the linear-to-linear polarization conversion efficiency to be more than 0.9. Upcoming CMB polarization experiments tend to cover the frequency range of $\delta \nu/\nu \sim 1$ or even broader in order to monitor the Galactic foreground (e.g. synchrotron and dust) emissions. Achromatic HWP (AHWP) is introduced by Pancharatnam~\cite{pancharatnam} and is employed by EBEX, the balloon-borne CMB experiment~\cite{ebex,matsumura_ahwp}. The stack of the multiple plates with properly chosen offset angles achieves the retardance of $\pi$ broader than $\delta \nu/\nu=0.3$. While this is a very attractive option, one complication with the Pancharatnam AHWP is that the amount of angle rotated by the AHWP becomes an electromagnetic frequency dependence. This effect can be rephrased as the polarization sensitivity axis depends on the instrument bandpass shape and the source spectrum. When the two or more sources are mixed, the uncertainty of the polarization sensitive angle is not only depending on the spectral shapes but also the relative polarized intensities. In this paper, we introduce the idea to mitigate the spectral dependence of the polarization angle. In section~2, we briefly review the AHWP polarimetery and introduce the mitigation recipe. In section~3, we show the experimental results as a demonstration of the idea. Finally in section~4 we discuss the actual implementations for forthcoming CMB experiments. \begin{figure}[tb] \centerline{\includegraphics[width=\columnwidth]{config.pdf}} \caption{Schematic view of the AHWP polarimeter (top) and ROC-AHWP polarimeter (bottom) using three wave plates ($m=3$).\label{fig:config.pdf}} \end{figure} | We introduced the recipe to mitigate the spectral dependent phase response of the AHWP polarimeter by introducing the second set of the AHWP that rotates or is stationary with respect to the first set of the AHWP. We experimentally show that the ROC-AHWP configuration with $m=3$ achieves $\Delta \phi_\nu < 1$~degrees while the phase of a single set of AHWP varies $\Delta \phi_\nu \sim 10$~degrees between 72 and 162~GHz. We also computationally show the potential ROC-AHWP design for POLARBEAR2 with $m=3$ and for LiteBIRD with $m=9$. Although we demonstrate this ROC-AHWP at the millimeter wavelength, the ROC-AHWP concept is applicable to any broadband polarimeter application. | 14 | 4 | 1404.5795 |
1404 | 1404.2660_arXiv.txt | \begin{description} \item[Background] Pulsar glitches---the sudden spin-up in the rotational frequency of a neutron star---suggest the existence of an angular-momentum reservoir confined to the inner crust of the neutron star. Large and regular glitches observed in the Vela pulsar have originally constrained the fraction of the stellar moment of inertia that must reside in the solid crust to about 1.4\%. However, crustal entrainment---which until very recently has been ignored---suggests that in order to account for the Vela glitches, the fraction of the moment of inertia residing in the crust must increase to about 7\%. This indicates that the required angular momentum reservoir may exceed that which is available in the crust. \item[Purpose] We explore the possibility that uncertainties in the equation of state provide enough flexibility for the construction of models that predict a large crustal thickness and consequently a large crustal moment of inertia. \item[Methods] Moments of inertia---both total and crustal---are computed in the slow-rotation approximation using a relativistic mean field formalism to generate the equation of state of neutron-star matter. \item[Results] We compute the fractional moment of inertia of neutron stars of various masses using a representative set of relativistic mean-field models. Given that analytic results suggest that the crustal moment of inertia is sensitive to the transition pressure at the crust-core interface, we tune the parameters of the model to maximize the transition pressure, while still providing an excellent description of nuclear observables. In this manner we are able to obtain fractional moments of inertia as large as 7\% for neutron stars with masses below 1.6 solar masses. \item[Conclusions] We find that uncertainties in the equation of state of neutron-rich matter are large enough to accommodate theoretical models that predict large crustal moments of inertia. In particular, we find that if the neutron-skin thickness of ${}^{208}$Pb falls within the (0.20-0.26)\,fm range, large enough transition pressures can be generated to explain the large Vela glitches---without invoking an additional angular-momentum reservoir beyond that confined to the solid crust. Our results suggest that \emph{the crust may be enough}. \end{description} | \label{Introduction} Rotation-powered pulsars tend to spin down slowly and steadily due to the emission of magnetic dipole radiation, making pulsars one of nature's most accurate clocks. For example, the Crab pulsar---with a rotational period of $P\! \approx \!33\,$ms---spins down at a rate of $\dot{P}/P\! \approx\!1.3\!\times\!10^{-11}\,{\rm s}^{-1}$ or about $13\,\mu{\rm s}$ per year. However, in spite of this remarkable regularity, young pulsars often display ``glitches'' which represent a sudden and abrupt spin-up in their rotational frequency. Through high-precision pulsar timing, an extensive glitch catalogue is now available which, at the time of this writing, has recorded a total of 439 glitches from 150 different pulsars\,\cite{Espinoza:2011pq,JBPGC}. Moreover, pulsar timing has revealed that glitches are recurrent---with some of the more active \emph{glitchers} being the Vela with 17, the Crab with 25, and PSR B1737-30 with 33\,\cite{JBPGC}. The glitch mechanism is intimately related to the formation of superfluid vortices in the inner crust of the rotating neutron star\,\citep[see Refs.][and references contained therein]{Anderson:1975zze,Pines:1985kz}. Given that many pulsars rotate extremely fast, the areal density of superfluid vortices may be extremely high. For example, in the case of the Crab pulsar the vortex density reaches $n_{\rm v}\!=\!4m_{n}/\hbar P\!\sim\!2\times\!10^{5}\,{\rm cm}^{-2}$, where $m_{n}$ is the neutron mass. The vortex density is so high that, although the bulk superfluid does not rotate, the superfluid as a whole appears to be rotating collectively as a ``rigid'' body\,\cite{Pines:1985kz}. However, as the pulsar slows down by the emission of magnetic dipole radiation, the initial distribution of vortices---which are believed to be pinned to a crystal lattice of neutron-rich nuclei---falls out of equilibrium. This induces a differential rotation between the slower neutron star and the faster superfluid vortices. When the differential lag is too large, then suddenly and abruptly some fraction of the vortices unpin, migrate outwards, and transfer their angular momentum to the solid crust---and to any stellar component strongly coupled to it (such as the liquid core). This sudden (typically in less than a day) transfer of angular momentum is detected as a pulsar spin-up or a \emph{glitch}. As a result of the glitch, the density of vortices diminishes from its pre-glitch value leaving the superfluid in close equilibrium with the solid crust. As the star continues to slow down over a period of days to years, stresses between the crust and the superfluid start to build up again until eventually more vortices unpin, transfer their angular momentum to the solid crust, and ultimately generate another glitch. The recurrence of glitches and the ensuing recovery is therefore a manifestation of a ``sluggish" solid crust falling periodically out of equilibrium with the larger distribution of superfluid vortices. Glitches from the Vela pulsar (B0833-45) were first observed in 1969 and have been continuously recorded for more than 40 years. So far a total of 17 glitches have been documented with individual spin-up rates of the order $\Delta\Omega/\Omega\!=\!10^{-6}$; note that although more glitches have been recorded for the Crab pulsar, the typical spin-up rate is three orders of magnitude smaller\,\cite{Espinoza:2011pq,JBPGC}. The high recurrence rate and large magnitude of the Vela glitches have been used to constrain the underlying equation of state (EOS) of neutron-star matter in terms of three measured quantities: its spin frequency $\Omega\!=\!70.338\,060\,018\,{\rm s}^{-1}$, its average spin-down rate $\dot{\Omega}\!=\!-9.8432\!\times\!10^{-11}$, and its \emph{glitch activity parameter} $A_{g}$---defined in terms of the cumulative spin-up rate $\sum_{n}\Delta\Omega_{n}/\Omega\!=\! 2.9760\!\times\!10^{-5}$ as follows\,\cite{Link:1999ca,Chamel:2012ae}: \begin{equation} A_{g} = \frac{1}{t}\sum_{n=1}^{N} \frac{\Delta\Omega_{n}}{\Omega} \approx 2.277 \!\times\!10^{-14}\,{\rm s}^{-1} \;, \label{VelaLimit1} \end{equation} where $t\!=\!41.421$ years is the total time elapsed between the first and the last ($N\!=\!17$) glitches. Thus, in the standard model of pulsar glitches, the ratio of the moment of inertia of the superfluid component driving the glitches $I_{s}$ to the moment of inertia of the solid crust $I_{c}$---\emph{plus any portion of the star strongly coupled to it}---must satisfy the following inequality\,\cite{Link:1999ca}: \begin{equation} \frac{I_{s}}{I_{c}} \ge A_{g}\frac{\Omega}{|\dot{\Omega}|} \approx 0.016. \label{VelaLimit2} \end{equation} That is, long term timing observations of the Vela pulsar suggest that in order to account for its glitch activity, at least 1.6\% of the stellar moment of inertia must reside in the superfluid reservoir\,\cite{Link:1999ca,Chamel:2012ae,Hooker:2013fda}. Moreover, by adopting two plausible assumptions, one may rewrite the above limit in such a way as to provide a more meaningful constraint on the EOS. First, in the two-component model of \citet{Baym:1969}, it is assumed that the component of the moment of inertia that rotates at the observed stellar frequency consists of the solid outer crust \emph{plus the liquid interior}; then, $I_{c}\!=\!I\!-\!I_{s}\!\approx\!I$, where $I$ is the \emph{total} stellar moment of inertia. Second, given that the inner crust is both thicker and denser than the outer crust, the crustal moment of inertia may be approximated as $I_{\rm crust}\!=\!I_{\rm outer}\!+\!I_{s}\!\approx\!I_{s}$. Thus, one may express the above constraint as follows: \begin{equation} \frac{I_{s}}{I_{c}} \simeq \frac{I_{\rm crust}}{I}\gtrsim\! 0.016 \,. \label{VelaLimit3} \end{equation} This expression is particularly convenient as both $I_{\rm crust}$ and $I$ may be readily evaluated in the \emph{slow-rotation approximation} (see Sec.\ref{Formalism}). Indeed, most of the formalism relies on the solution of the Tolman-Oppenheimer-Volkoff equation whose only required input is the EOS of neutron-star matter. Further, whereas the total moment of inertia depends sensitively on the poorly constrained high-density component of the EOS, the crustal component $I_{\rm crust}$ is sensitive to physics that may be probed in the laboratory. In fact, analytic expressions for $I_{\rm crust}$ already exist that are both illuminating and highly accurate\,\citep[see Refs.][and references contained therein]{Link:1999ca,Lattimer:2006xb,Fattoyev:2010tb}. In particular, the crustal moment of inertia is highly sensitive to physical observables---particularly the transition pressure---at the crust-core interface. Recently, however, the standard glitch mechanism has been called into question\,\cite{Andersson:2012iu,Chamel:2012ae}. It has been argued that \emph{crustal entrainment}---the non-dissipative elastic scattering of unbound neutrons by the crystal lattice---effectively reduces the angular-momentum reservoir. That is, entrainment effects reduce the density of superfluid neutrons that could eventually become the source of the superfluid vortices. The impact from crustal entrainment may be encoded in an effective neutron mass $m_{n}^{\star}\!\equiv\!m_{n}n_{n}^{\rm f}/n_{n}^{\rm c}$, that reflects the ratio of unbound neutrons $n_{n}^{\rm f}$ ({\sl i.e.,} neutrons not bound to the nuclear lattice) to those $n_{n}^{\rm c}$ neutrons that are effectively free ({\sl i.e.,} not entrained)\,\cite{Chamel:2011aa,Chamel:2012ae}. As a result, crustal entrainment modifies the constraint given in Eq.\,(\ref{VelaLimit3}) to \begin{equation} \frac{I_{\rm crust}}{I}\gtrsim\!0.016\left( \frac{\langle m_{n}^{\star}\rangle}{m_{n}}\right)\simeq 0.07\;, \label{VelaLimit4} \end{equation} where in the last expression we have adopted $\langle m_{n}^{\star}\rangle/m_{n}\!\simeq\!4.3$ as suggested in Ref.\,\cite{Andersson:2012iu}. Note that the amount of crustal entrainment may be uncertain, so more sophisticated calculations could yield a different constraint for the fractional moment of inertia. Here we assume, as in Ref.\,\cite{Andersson:2012iu}, that its value lies between the two limits given in Eqs.\,(\ref{VelaLimit3}) and \,(\ref{VelaLimit4}); that is, $0.016\!\lesssim I_{\rm crust}/I\!\lesssim\,0.07$. Given the significant impact that crustal entrainment may have in constraining $I_{\rm crust}/I$, it has been suggested that \emph{the crust is not enough}, so that the core superfluid must also participate in glitches\,\cite{Andersson:2012iu}. However, before completely dismissing the standard glitch mechanism in favor of a more exotic explanation, we explore the conservative alternative that uncertainties in the EOS (which are large) could still allow the inner crust to account---by itself---for a large enough fraction of the moment of inertia to explain the large Vela glitches. Uncertainties in the EOS are known to significantly affect the transition between the liquid interior and the solid crust. Particularly relevant to the present discussion is the poorly-known density dependence of the symmetry energy $S(\rho)$. The symmetry energy represents the energy required to convert symmetric nuclear matter into pure neutron matter at a fixed baryon density $\rho$. Although the symmetry energy at saturation density is fairly well determined, the value of its slope $L$ remains highly uncertain. And it is precisely the slope of the symmetry energy that controls the properties of neutron-rich matter at the crust-core transition region\,\cite{Horowitz:2000xj}. Indeed, the transition density and proton fraction at the crust-core interface are both linearly anti-correlated to $L$. However, in stark contrast, the transition pressure does not vary monotonically with $L$\,\cite{Ducoin:2010as,Fattoyev:2010tb}. Note that such lack of correlation between the transition pressure and $L$---which appears to contradict some recent studies\,\cite{Worley:2008cb, Xu:2008vz,Xu:2009vi,Moustakidis:2010zx,Paar:2014qda}---emerges only as one systematically explores a wide range of models. Given that the crustal moment of inertia grows with an increasing transition pressure, we want to explore the possibility of generating realistic EOS that predict large transition pressures. As we shall demonstrate, we find a class of relativistic mean field models with moderate values of $L$ that generate fractional moments of inertia as large as $I_{\rm crust}/I\!\approx\!0.09$ for a canonical $1.4\,M_{\odot}$ neutron star. By using neutron star observations, Ref.\,\cite{Steiner:2014pda} has recently and independently reached a similar conclusion. Moreover, we argue that these predictions may be directly tested in the laboratory. Although $L$ can not be directly measured in the laboratory, it is known to be strongly correlated to the neutron-skin thickness of ${}^{208}$Pb\,\cite{Brown:2000, Furnstahl:2001un,Centelles:2008vu,RocaMaza:2011pm}---a fundamental nuclear-structure observable that has been measured\,\cite{Abrahamyan:2012gp,Horowitz:2012tj}---and will be measured with increasing accuracy---at the Jefferson Laboratory\,\cite{PREXII:2012}. In summary, it is the main goal of the present work to explore whether present uncertainties in the EOS are large enough to accommodate realistic models that could account for large pulsar glitches---even in the presence of crustal entrainment. We have organized the paper as follows. In Sec.\,\ref{Formalism} we review the essential details required to compute the stellar moment of inertia and the class of equations of state that will be used. In particular, special attention is paid to the crustal component of the moment of inertia and its sensitivity to the transition pressure at the crust-core interface. In Sec.\,\ref{Results} we provide predictions for the fraction of the moment of inertia residing in the crust. We show that although the transition pressure is sensitive to the slope of the symmetry energy, its behavior is not monotonic. This suggests a range of values for $L$---that are neither too small nor too large---that provide the thickest crust and consequently the largest fractional moment of inertia. Finally, we offer our conclusions in Sec.\,\ref{Conclusions}. | \label{Conclusions} The large jumps observed in the spin frequency of some neutron stars like the Vela pulsar suggest the existence of an angular-momentum reservoir confined to the inner stellar crust. As such, large pulsar glitches have been used to constrain the EOS of neutron-rich matter by imposing limits on the fraction of the moment of inertia that must reside in the crust. Until recently, the large Vela glitches demanded that at least 1.6\% of the stellar moment of inertia must be located in the crust. Although useful, many realistic EOS are compatible with such limit. However, recent studies seem to indicate that a significantly larger value is required once crustal entrainment is taken into account. Indeed, encoding the impact of crustal entrainment in a value of the effective neutron mass of $\langle m_{n}^{\ast}\rangle/m_{n}\!=\!4.3$, suggests that the limit on the fractional moment of inertia must be increased to almost 7\%. Unless the mass of the Vela pulsar is very small ($\lesssim\!1\,M_{\odot}$) then a large number of theoretical models become incompatible with this much more stringent limit. This has lead to the assertion that the ``crust is not enough". The main goal of this work was to re-examine whether indeed the crust is not enough. Given that the crustal moment of inertia is highly sensitive to the transition pressure $P_{t}$ at the crust-core interface, we examined the predictions of a variety of relativistic mean-field models in the transition region. In particular, we found that certain bulk properties of the EOS at the crust-core interface are strongly correlated to the slope of the symmetry energy $L$. However, not the transition pressure. Indeed, $P_{t}$ increases monotonically for small $L$, reaches its maximum at some intermediate value, and ultimately decreases with increasing $L$. Given that in the class of RMF models used in this work one can easily tune the value of $L$, we searched for models with the largest transition pressure. By doing so, we were able to generate neutron stars with thick crusts and large crustal moments of inertia. Indeed, our results support the standard model of pulsar glitches---provided the Vela mass does not exceed $1.6\,M_{\odot}$. In particular, this requires values for the neutron-skin thickness of ${}^{208}$Pb of about $R_{\rm skin}^{208}\!\simeq\!(0.20\!-\!0.26)\,{\rm fm}$. This finding---which represents the central result of our work---offers yet another attractive connection between a laboratory measurement and an astrophysical observation. We close with a few questions and comments on the impact of our findings on other neutron-star observables sensitive to crustal dynamics. First, rapidly rotating neutron stars with large asymmetries (such as mass quadrupoles) are efficient sources of gravitational waves. In particular, ``large mountains'' on rapidly rotating neutron stars may efficiently radiate gravitational waves provided that the breaking strain of the crust is large\,\cite{Andersson:2009yt}. Horowitz and Kadau have performed large-scale Molecular-Dynamics simulations that reveal a large breaking strain in support of large mountains in neutron stars\,\cite{Horowitz:2009ya}. Such findings suggest that large mountains in rapidly rotating neutron stars may generate gravitational waves that may be detectable by the next generation of gravitational-wave observatories. Thus, one would like to understand the impact of the large crustal thicknesses found in our work ({\sl e.g.,} $R_{\rm crust}\!\simeq\!2.4\,{\rm km}$ for a $1.4\,M_{\odot}$ neutron star) on the breaking strain of the crust and ultimately on the emission of gravitational waves. Moreover, the size of the crust may also impact the cooling light curves of low-mass X-ray binaries during quiescence. Indeed, a thick stellar crust could increase the time scale for crustal cooling after extended periods of accretion; see Ref.\,\cite{Brown:2009kw} and references contained therein. However, although the large crustal thicknesses raises interesting questions and accounts for large crustal fractions of the stellar moment of inertia, the moderate values of $L$ required to account for the large pulsar glitches are inconsistent with a recent analysis of quiescent low mass X-ray binaries that suggests very small stellar radii. In fact, the class of relativistic mean-field models employed in this work are unable to generate such small radii---regardless of whether one incorporates the pulsar-glitch constraint or not. Although not without controversy, we found instructive to take the small-radius result at face value and ask whether it is possible to account for both large pulsar glitches and small neutron stars. If crustal entrainment is as large as it has been suggested, then a moderate value of the neutron-skin thickness of ${}^{208}$Pb is required. This prediction can be tested by performing parity violating electron scattering experiments at the Jefferson Laboratory. However, in order for neutron-star radii to be as small as suggested, a dramatic softening of the symmetry energy must develop by the time that the density reaches about 2 times nuclear-matter saturation density. Such a rapid softening is likely to involve a change in the structure of dense matter---likely due to a phase transition. In principle, the onset of such a phase transition may also be probed in the laboratory using energetic heavy-ion collisions of highly-asymmetric nuclei. Finally, the EOS must significantly stiffen at even higher densities in order to account for the observation of massive neutron stars. Such unique behavior of the EOS and its possible realization in neutron stars reaffirms the special bond between astrophysics and nuclear physics. | 14 | 4 | 1404.2660 |
1404 | 1404.5937_arXiv.txt | We present the results of a study of the star HD\,34736. The spectropolarimetric observations carried out at the 6-m telescope showed the presence of a strong variable longitudinal magnetic field, exceeding $-4500$~G. The analysis of the HIPPARCOS photometry gives a set of possible periods of the brightness variability of the star, of which $0\fd3603$ is preferred. The variable radial velocity of spectral lines of the star and some signatures of lines of at least one other component show that HD\,34736 is a double short-period system. Modeling of the spectra allowed us to estimate the effective temperature $T_\mathrm{eff}$~of the stars ($13\,700$ and $11\,500$~K) and their projected rotational velocities $v\sin i$ ($73$ and $\geq90$~km\,s$^{-1}$). The analysis of all the available information about the star allows us to hypothesize that the object of study is a close, possibly interacting binary system. | The study of magnetic properties of early-type stars provides a unique opportunity to study in detail the features of the processes of their evolution, formation and variation of the magnetic field strength with age. This is primarily due to the rapid evolution of massive stars. A major part of B-type stars is found in open clusters and associations, and hence the accuracy of determination of their age is significantly higher than for the field stars. All these factors have facilitated the start of a large observational program called {\it Magnetic Fields of Massive Stars} at the Special Astrophysical Observatory of Russian Academy of Sciences (SAO RAS). The conceptual issues of the program, formulating the problem, and selecting the objects of study are presented in the paper by Romanyuk and Yakunin~\cite{romanyuk_yakunin:2012}. Spectropolarimetric observations with the 6-m BTA telescope within this program started at the end of 2010~and are still ongoing. Recently, our research has been focused on the study of the already known and the search for new stars with magnetic fields in the nearby Ori\,OB1 association. In the course of the survey of chemically peculiar (CP) stars of the association, the star HD\,35298 was studied in detail~\cite{yakunin:2013}. The observations of poorly-studied CP stars of the association allowed us to detect the signs of magnetic field in four stars. This paper is devoted to the features of one of them, HD\,34736. A chemically peculiar star HD\,34736 ($m_V=7\fm82$) with silicon abundance anomalies~\cite{roman:1978} is a member of the Ori\,OB1 association (sub-group~C)~\cite{brown_etal:1994}. A large amount of the photometric data (the systems $UBV$~\cite{deutschman_etal:1976}, Str\"omgren~\cite{vogt_faundez:1979}, Maitzen~$\Delta a$~\cite{maitzen_vogt:1983}, Walraven~\cite{de-geus_etal:1990}) provides confident estimates of the main characteristics of the star. According to the Geneva photometry indices, North and Kramer~\cite{north_cramer:1984} predict the surface magnetic field of the star of ~\mbox{$B_\mathrm{s}\approx1.9$~kG.} However, the literature contains no evidence on the direct measurement of this magnetic field. Based on the analysis of data presented in the work of Brown et al.~\cite{brown_etal:1994}, the catalog of CP stars by Renson and Manfroid~\cite{renson_manfroid:2009}, and other literary sources, we have selected the star HD\,34736 as a candidate for the spectropolarimetric observations with the 6-m telescope of the SAO RAS~\cite{romanyuk_etal:2013}. Three observations of the star were performed to date, their description is contained in Section~2. The results of measurements of the Zeeman effect in the spectra of the star and determining the magnitude of its magnetic field are shown in Section~3. Section 4 is devoted to the estimation of the main physical parameters of HD\,34736. The final Section 5 discusses the results we have obtained. | The new magnetic star we have discovered, HD\,34736, is a unique object. Three measurements of polarized spectra show a variable magnetic field, the magnitude of the longitudinal component of which varies from 0 to almost \mbox{$-4500$}~G. Along with the magnetic field we also observe a spectral variability. The character of the latter indicates the presence in the spectrum of the lines belonging to the second star, somewhat cooler \mbox{($T_\mathrm{eff}=11500$~K)}. The study of the HIPPARCOS photometry data allows to select several possible periods of brightness variability, where the most likely of them is $0\fd3603$. Apparently, this value is the orbital period of the binary system. Radial velocities of the components, obtained by decomposing the spectrum into components, support this assumption. All our results concerning the modeling of the spectrum of the second star are preliminary. For a more detailed analysis new observations are required in order to measure the radial velocity of the system and the longitudinal magnetic field of the star. The rotation period of the magnetic star in the HD\,34736 system can only be determined after the analysis of a sufficient number of measurements of its magnetic field. To make the final conclusion on the presence of a magnetic field in the second star, circularly polarized spectra with a very high signal-to-noise ratio are required. We have tried to divide the spectrum of the star into the components which allowed us to estimate the effective temperatures, surface gravity, and rotation velocities of both components. We could obtain the preliminary value of \mbox{$v\sin i=130$ km\,s$^{-1}$} for the secondary component only from one line of Mg\,II~4481\,\AA, which manifests itself well in the right wing of the line of the composite spectrum from \mbox{October 23, 2013.} If we assume that the main contribution to the observed spectrum is made by two stars with effective temperatures of $T_\mathrm{eff}$ equal to $13\,700$ and $11\,500$~K, a larger part of the observed hydrogen H$_\beta$ line can be described well, except for its short-wavelength wing. At the same time, the shapes of many lines reveal features that can be interpreted as traces of the third spectrum with the chemical composition and temperature of the main star. The nature of this spectrum remains to be clarified, but we can now confidently say that its manifestation is not a result of processing errors: the short-wave wing of the composite line H$_\beta$ can be described with high accuracy only in the presence of the third spectrum. In our opinion the physical explanation of such a complex spectrum may be as follows. The star HD\,34736 represents a short-period binary system, the main component of which, with a higher temperature, is a magnetic star. The stars make one revolution around the common center of gravity in about $0\fd3603$. This results in substantially different radial velocities of individual components. The brightness curve of HD\,34736 (Fig.~\ref{figure:2}) has a second minimum at phase \mbox{$\varphi\approx0.73$} which can be caused by the spatial orientation of the system relative to the observer and the dimension ratio of stars. It is very likely that interchange of matter takes place in this system, forming a shell or a tail. We detect the traces of this matter in form of the third component of the spectrum. The presence in it of the same lines as in the spectrum of the main star can tell of the origin of matter of this shell. The spectral range we observe bears no prominent signs of emission. However, Leone~\cite{leone:1994} states that HD\,34736 is a confirmed X-ray source. The X-ray activity in the case of main-sequence B-type stars is often a sign of either the presence of a powerful magnetosphere or close interacting components. It is possible that HD\,34736 combines both cases. Therefore, the further study of this star can be extremely important to address the origin and evolution of close binary systems possessing a strong magnetic field. But at the same time, we cannot exclude other possible explanations of observations, since the total data set is still little. {\bf Acknowledgments}\\ This work was partially supported by the Russian Foundation for Basic Research (grant no.~\mbox{12-02-00009-a}). The observations on the \mbox{6-meter} BTA telescope were conducted with the financial support of the Ministry of Education and Science of the Russian Federation (state contracts no.~14.518.11.7070, 16.518.11.7073). | 14 | 4 | 1404.5937 |
1404 | 1404.7103_arXiv.txt | We present thermal Sunyaev-Zel'dovich effect (SZE) measurements for 42 galaxy clusters observed at 150 GHz with the APEX-SZ experiment. For each cluster, we model the pressure profile and calculate the integrated Comptonization $Y$ to estimate the total thermal energy of the intracluster medium (ICM). We compare the measured $Y$ values to X-ray observables of the ICM from the literature (cluster gas mass $M_{\rm{gas}}$, temperature $T_X$, and $Y_X =M_{\rm{gas}}T_X$) that relate to total cluster mass. We measure power law scaling relations, including an intrinsic scatter, between the SZE and X-ray observables for three subsamples within the set of 42 clusters that have uniform X-ray analysis in the literature. We observe that differences between these X-ray analyses introduce significant variability into the measured scaling relations, particularly affecting the normalization. For all three subsamples, we find results consistent with a self-similar model of cluster evolution dominated by gravitational effects. Comparing to predictions from numerical simulations, these scaling relations prefer models that include cooling and feedback in the ICM. Lastly, we measure an intrinsic scatter of $\sim28$ per cent in the $Y-Y_X\,$ scaling relation for all three subsamples. | \label{SEC:introduction} \setcounter{footnote}{0} As the largest gravitationally collapsed objects in the Universe, clusters of galaxies provide a unique opportunity to study the evolution of large-scale structure. The distribution and abundance of clusters is sensitive to both the geometry of the universe and the growth of density perturbations \citep[e.g.,][]{haiman2001,weller2002}. Currently, cluster-based constraints on cosmology are limited by systematic uncertainties in relating observables to cluster masses. Most of the cluster mass is in the form of dark matter and therefore is not directly observable. Instead, cluster masses are inferred through scaling relations with observable signals such as X-ray luminosity, galaxy velocity distribution, weak-lensing shear and Sunyaev-Zel'dovich effect (SZE) brightness. Under the model of self-similarity, where cluster evolution is dominated by gravitational processes, cluster mass scales with observable signals through simple power law relations \citep{kaiser1986}. In this model, the intracluster medium (ICM) is in hydrostatic equilibrium, for which the scaling to cluster mass can be predicted for a given observable. However, self-similarity does not take into account the role of non-thermal mechanisms such as turbulent gas motions in estimating cluster mass. This simple model also neglects the effects of physical processes internal to the cluster such as feedback from active galactic nuclei and star formation. Numerical simulations predict that while the power law exponent of SZE scaling relations will be consistent with self-similarity, the normalization does depend on the internal cluster astrophysics \citep{motl2005, nagai2006, lau2009}. An empirical measurement of the scaling relations therefore informs models of cluster astrophysics, progressing towards the needed calibration for cosmology. Currently, cluster surveys are operating in the optical \citep[e.g.,][]{gilbank2011}, X-ray \citep[e.g.,][]{vikhlinin2009,mantz2010}, and the millimeter \citep[e.g.,][]{carlstrom2011,swetz2011,planck2013xxix} wavelength regimes. However, optical and X-ray measurements of clusters suffer from cosmological dimming, and only the brightest and most massive clusters are detected at high redshifts ($z > 1$). In contrast, the millimeter-wavelength thermal SZE \citep{sunyaev72}, where Cosmic Microwave Background (CMB) photons inverse-Compton scatter off hot intracluster electrons, is redshift-independent. SZE surveys with sufficient resolution to resolve clusters, such as those performed with the South Pole Telescope \citep{reichardt2013} and the Atacama Cosmology Telescope \citep{hasselfield2013} detect clusters with a mass selection nearly independent of redshift. Precision cosmology requires that mass-observable scaling relations be characterized with a high level of accuracy, including both the measurement uncertainty and the intrinsic scatter from cluster-to-cluster differences. Numerical simulations suggest that SZE observations are relatively insensitive to the details of cluster astrophysics \citep[e.g.,][]{hallman2006, nagai2006}, resulting in low intrinsic scatter scaling relations and a tighter constraint on cluster mass. Previous measurements of SZE scaling relations include interferometric observations from the OVRO/BIMA \citep{bonamente2008} and SZA \citep{culverhouse2010, marrone2012} arrays as well as imaging studies with the South Pole Telescope \citep{andersson2011,plagge2010,benson2013}, BOLOCAM \citep{sayers2011}, Atacama Cosmology Telescope \citep{marriage2010}, and the Planck mission \citep{planck2011XI}. In general, those studies find that observational SZE scaling relations agree with expectations based on self-similarity and that simulations including additional non-adiabatic physics are preferred. While cluster samples selected from large surveys (both SZE and X-ray) have well-known selection functions, studies such as the one presented in this paper that target known clusters often select them in an \textit{ad hoc} manner. The influence of this sample selection on SZE scaling relations is unknown. In this paper, we study how the SZE signal scales with X-ray observables related to cluster mass using observations from the APEX-SZ imaging bolometer array. We measure SZE scaling relations using X-ray observables ($Y_X$, $M_{\rm{gas}}$, and $T_X$) as proxies for the total cluster mass. APEX-SZ observed a small set of 11 clusters selected from the REFLEX X-ray survey (referred to as the REFLEX-DXL sample \citep{zhang2006}) as well as an additional 31 clusters selected in an \textit{ad hoc} manner. Within this full sample are clusters drawn from the \citet[][hereafter Z08]{zhang2008} and \citet[][hereafter M10]{mantz2010} samples. We measure power law scaling relations for each of these three subsamples (REFLEX-DXL, Z08, M10) and compare the results to the expectations of the self-similar model as well as numerical simulations that incorporate different physical processes in the ICM. We compare the results between the three cluster samples as well as to the full APEX-SZ sample to investigate the effects of sample selection and uniform analysis on the measured power law and intrinsic scatter of the SZE scaling relations. In this paper, we assume the \textit{WMAP7}+BAO+$H_0$ $\Lambda$CDM best-fit cosmology in which $H_0=70.4$ km\,$\rm{s}^{-1}$\,$\rm{Mpc}^{-1}$, $\Omega_M=0.272$, and $\Omega_{\Lambda}=0.728$ \citep{komatsu2011}. The structure of the paper is as follows. Section~\ref{sec:theory} reviews the SZE and the associated scaling relations. Section \ref{SEC:observations} introduces the clusters in this study and observations thereof. Section \ref{SEC:szdatareduction} briefly describes the data reduction process for the APEX-SZ instrument. We present the SZE-X-ray scaling relations and discuss these results in Section~\ref{SEC:results}. A summary and an outlook on future work are provided in Section \ref{SEC:conclusions}. | \label{SEC:conclusions} We present Sunyaev-Zel'dovich effect (SZE) observations of galaxy clusters measured with the APEX-SZ experiment and use them to investigate the scaling of the SZE with cluster mass. We model the thermal pressure of the intracluster medium using the \citet{arnaud2010} universal pressure profile and use the results to calculate the spherical integrated Comptonization \y500 for each cluster. We compare these measurements of \y500 to X-ray estimates of cluster mass taken from the literature. Using these two sets of observables, we measure the \yyx, \ymgas, and \ytx scaling relations, finding the best-fit power law and intrinsic scatter for each. These scaling relations are measured for three subsamples of the 42 APEX-SZ clusters that have uniform X-ray analysis: the REFLEX-DXL sample, 15 clusters from Z08, and 19 clusters from M10. We find when all 42 clusters with varying X-ray analyses are included, significant systematics are introduced into the power law regression parameters and the intrinsic scatter increases. For all three subsamples, we find that the best-fit power laws for the \yyx, \ymgas, and \ytx relations have exponents consistent with those predicted by the self-similar model. We compare the measured normalizations for each scaling relation to numerical simulations to probe the underlying astrophysics of the intracluster medium (ICM). The measured normalization of the \yyx relation for two of the subsamples implies a higher $Y_{500}d_A^2/Y_X$ than seen in previous studies. However, there is a large variation in the normalization between the three samples, despite the uniformly analyzed X-ray data. Therefore, we associate a large systematic uncertainty with the high $Y_{500}d_A^2/Y_X$ and do not draw further conclusions from it. We compare the normalization for the \ymgas and \ytx relations to the numerical simulations of \citet{nagai2006} and find a weak preference for models that included radiative cooling and feedback in the ICM as well as standard gas dynamics. Finally, we find that the levels of intrinsic scatter for the \yyx and \ymgas relations are consistent with previous measurements. The uncertainties in the intrinsic scatter are large due to the small number of clusters in the three subsamples. A larger sample is key to improving this measurement of intrinsic scatter in \y500 and its application as an estimator for total cluster mass in cluster-based cosmological constraints. A uniform X-ray analysis for the full sample of APEX-SZ clusters is ongoing. This analysis will resolve the main systematic limitation of the full cluster sample and will allow us to use its large number of clusters to improve on the constraints presented here for the three subsamples. Additionally, a follow-up program of optical observations has been concluded to estimate the total cluster mass independently using weak-lensing measurements. Future papers will use the additional X-ray and weak-lensing information to improve further our understanding of how \y500 scales with total cluster mass and facilitate its use in exploring the physics of the ICM as well as in constraining cosmological parameters using galaxy clusters. | 14 | 4 | 1404.7103 |
1404 | 1404.1455_arXiv.txt | We present the first results of searches for axions and axion-like-particles with the \Xehund\ experiment. The axion-electron coupling constant, \gAe, has been probed by exploiting the axio-electric effect in liquid xenon. A profile likelihood analysis of 224.6 live days $\times$ 34 kg exposure has shown no evidence for a signal. By rejecting \gAe\, larger than $7.7 \times 10^{-12}$ (90\% CL) in the solar axion search, we set the best limit to date on this coupling. In the frame of the DFSZ and KSVZ models, we exclude QCD axions heavier than 0.3 eV/c$^2$ and 80 eV/c$^2$, respectively. For axion-like-particles, under the assumption that they constitute the whole abundance of dark matter in our galaxy, we constrain \gAe\, to be lower than $1 \times 10^{-12}$ (90\% CL) for masses between 5 and 10 keV/c$^2$. | Axions were introduced in the Peccei-Quinn solution of the strong CP problem as pseudo-Nambu-Goldstone bosons emerging from the breaking of a global U(1) symmetry~\cite{pecceiquinn1977,weinberg1978, wilczeck1978}. Although this original model has been ruled out, ``invisible'' axions arising from a higher symmetry-breaking energy scale are still allowed, as described, for example, in the DFSZ and KSVZ models~\cite{DFS, Z, K, SVZ}. In addition to QCD axions, axion-like particles (ALPs) are pseudoscalars that do not necessarily solve the strong CP problem, but which have been introduced by many extensions of the Standard Model of particle physics. Axions as well as ALPs are well motivated cold dark matter candidates~\cite{Abbott:1982af}. Astrophysical observations are thought to be the most sensitive technique for detecting axions and ALPs~\cite{sikivie1983}: the Sun would constitute an intense source of this particles (referred to as solar axions), where they can be produced via Bremsstrahlung, Compton scattering, axio-recombination and axio-deexcitation~\cite{redondo}. Additionally, searches can be conducted for ALPs that may have been generated via a non-thermal production mechanism in the early universe and which now constitute the dark matter in our galaxy (referred to as galactic ALPs). \\ Axions and ALPs may give rise to observable signatures in detectors through their coupling to photons (\gAg), electrons (\gAe) and nuclei (\gAN). The coupling \gAe\, may be tested via scattering off the electron of a target, such as liquid xenon (LXe), through the axio-electric effect~\cite{dimopoulos1986, avignone1987, pospelov2008, derevianko2010, arisaka2013}. This process is the analogue of the photo-electric process with the absorption of an axion instead of a photon. We report on the first axion searches performed with the \Xehund\, experiment. The expected interaction rate is obtained by the convolution of the flux and the axio-electric cross section. The latter is given, both for QCD axions and ALPs, by \begin{equation} \sigma_{Ae} = \sigma_{pe}(E_A)\frac{{g_{Ae}}^2}{\beta_A}{\frac{3{E_A}^2}{16\pi\,\alpha_{em}\,{m_e}^2}}\left(1-\frac{\beta_A^{2/3}}{3} \right), \label{eq:sigmaA} \end{equation} as described in~\cite{avignone1987, pospelov2008, derevianko2010, arisaka2013, cuore2013}. In Eq.(\ref{eq:sigmaA}), $\sigma_{pe}$ is the photoelectric cross section for LXe~\cite{xpe}, $E_A$ is the axion energy, $\alpha_{em}$ is the fine structure constant, $m_e$ is the electron mass, and $\beta_A$ is the axion velocity over the speed of light, $c$. The solar axion flux has recently been recalculated in~\cite{redondo}. This incorporates four production mechanisms that depend upon $g_{Ae}$: Bremsstrahlung, Compton scattering, atomic recombination, and atomic deexcitation. The corresponding flux is 30\% larger than previous estimates due to atomic recombination and deexcitation, which previously were not taken into account. However, \cite{redondo} does not include corrections for axions heavier than 1 \keVcc, which we therefore takes as an upper mass limit for our analysis. For solar axions, both flux and cross-section depend upon $g_{Ae}^2$, thus the interaction rate scales with the fourth power of the coupling. For non-relativistic ALPs in the galaxy, assuming that they constitute the whole dark matter halo density ($\rho_{DM} \sim 0.3$ GeV/cm$^3$ \cite{Green:2011bv}), the total flux is given by $\phi_\text{ALP} = c \beta_A \times \rho_{DM} / m_A$, where $m_A$ is the ALP mass. The interaction rate for these ALPs depends on $g_{Ae}^2$, as the flux is independent from the axion coupling. As $\beta_A \approx 10^{-3}$\, in the non-relativistic regime, the velocities cancel out in the convolution between $\sigma_{Ae}$ and the flux. Thus the expected electron recoil spectrum is independent from the particle speed. As the kinetic energy of the ALPs is negligible with respect to their rest mass, a monoenergetic peak at the axion mass is expected in the spectrum. | 14 | 4 | 1404.1455 |
|
0710 | 0710.5082.txt | % context heading (optional) % {} leave it empty if necessary {} % aims heading (mandatory) {Photoionization models so far are unable to account for the high electron temperature $T_e$([\ion{O}{iii}]) implied by the line intensity ratio [\ion{O}{iii}]$\lambda$4363\AA/[\ion{O}{iii}]$\lambda$5007\AA\ in low-metallicity blue compact dwarf galaxies, casting doubts on the assumption of photoionization by hot stars as the dominant source of heating of the gas in these objects of large cosmological significance.} % methods heading (mandatory) {Combinations of runs of the 1-D photoionization code NEBU are used to explore alternative models for the prototype giant \ion{H}{ii} region shell I\,Zw\,18\,NW, with no reference to the filling factor concept and with due consideration for geometrical and stellar evolution constraints.} % results heading (mandatory) {Acceptable models for I\,Zw\,18\,NW are obtained, which represent schematically an incomplete shell comprising radiation-bounded condensations embedded in a low-density matter-bounded diffuse medium. The thermal pressure contrast between gas components is about a factor 7. The diffuse phase can be in pressure balance with the hot superbubble fed by mechanical energy from the inner massive star cluster. The failure of previous modellings is ascribed to (1) the adoption of an inadequate small-scale gas density distribution, which proves critical when the collisional excitation of hydrogen contributes significantly to the cooling of the gas, and possibly (2) a too restrictive implementation of Wolf-Rayet stars in synthetic stellar cluster spectral energy distributions. A neutral gas component heated by soft X-rays, whose power is less than 1\% of the star cluster luminosity and consistent with CHANDRA data, can explain the low-ionization fine-structure lines detected by SPITZER. [O/Fe] is slightly smaller in I\,Zw\,18\,NW than in Galactic Halo stars of similar metallicity and [C/O] is correlatively large.} % conclusions heading (optional), leave it empty if necessary {Extra heating by, \eg, dissipation of mechanical energy is not required to explain $T_e$([\ion{O}{iii}]) in I\,Zw\,18. Important astrophysical developments are at stakes in the 5\% uncertainty attached to \oiii\ collision strengths.} | \label{intro} The optical properties of Blue Compact Dwarf (BCD) galaxies are similar to those of Giant Extragalactic \hii\ Regions (GEHIIR). Their blue continuum arises from one or several young Massive Star Clusters (MSC), which harbour extremely large numbers of massive stars. BCDs are relatively isolated, small-sized, metal-poor galaxies (Kunth \& \"Ostlin 2000) and may be the rare `living fossils' of a formerly common population. BCDs can provide invaluable pieces of information about the primordial abundance of helium (\eg, Davidson \& Kinman 1985), the chemical composition of the InterStellar Medium (ISM, \eg, Izotov et al. 2006), the formation and evolution of massive stars, and the early evolution of galaxies at large redshift. Among them, \IZ\ stands out as one of the most oxygen-poor BCDs known (\eg, Izotov et al. 1999) and a young galaxy candidate in the Local Universe (\eg, Izotov \& Thuan 2004). The line emission of \hii\ regions is believed to be governed by radiation from massive stars, but spectroscopic diagnostics most often indicate spatial fluctuations of the electron temperature \Te\ (see the dimensionless parameter $t^2$, Peimbert, 1967), that appear larger than those computed in {\sl usual} photoionization models, suggesting an {\sl extra heating} of the emitting gas (\eg, Peimbert, 1995; Luridiana et al. 1999). Until the cause(s) of this failure of photoionization models can be identified, a sword of Damocles is hung over a basic tool of astrophysics. Tsamis \& P\'equignot (2005) showed that, in the GEHIIR 30\,Dor of the LMC, the various \Te\ diagnostics could be made compatible with one another if the ionized gas were {\sl chemically inhomogeneous} over small spatial scales. A pure photoionization model could then account for the spectrum of a bright filament of this nebula. Although this new model needs confirmation, it is in suggestive agreement with a scenario by Tenorio-Tagle (1996) of a recycling of supernova ejecta through a rain of metal-rich droplets cooling and condensing in the Galaxy halo, then falling back on to the Galactic disc and incorporating into the ISM without significant mixing until a new \hii\ region eventually forms. If this class of photoionization models is finally accepted, extra heating will not be required for objects like 30\,Dor, with near Galactic metallicity. Another problem is encountered in low-metallicity (`low-Z') BCDs (Appendix~A). In BCDs, available spectroscopic data do not provide signatures for $t^2$'s, but a major concern of photoionization models is explaining the high temperature \Te(\oiii) infered from the observed intensity ratio $r$(\oiii) = \oiii\la4363/(\oiii\la5007+4959). Thus, Stasi\'nska \& Schaerer (1999, SS99) conclude that photoionization by stars fails to explain $r$(\oiii) in the GEHIIR \IZ\,NW and that photoionization must be supplemented by other heating mechanisms. A requirement for extra heating is indirectly stated by Luridiana et al. (1999) for NGC\,2363. A possible heating mechanism is conversion of mechanical energy provided by stellar winds and supernovae, although a conclusion of Luridiana et al. (2001) does not invite to optimism. A limitation of this mechanism is that most of this mechanical energy is likely to dissipate in hot, steadily expanding superbubbles (Martin, 1996; Tenorio-Tagle et al. 2006). It is doubtful that heat conduction from this coronal gas could induce enough localized enhancement of \Te\ in the photoionized gas (\eg, Maciejewski et al. 1996), even though Slavin et al. (1993) suggest that turbulent mixing may favour an energy transfer. Martin (1997) suggests that shocks could help to explain the trend of ionization throughout the diffuse interstellar gas of BCDs, but concedes that ``shocks are only being invoked as a secondary signal in gas with very low surface brightness''. Finally, photoelectric heating from dust is inefficient in metal-poor hot gas conditions (Bakes \& Tielens 1994). Nevertheless, the conclusion of SS99 is now accepted in many studies of GEHIIRs. It entails so far-reaching consequences concerning the physics of galaxies at large redshifts as to deserve close scrutiny. If, for exemple, the difference between observed and computed \Te(\oiii) in the model by SS99 were to be accounted for by artificially raising the heat input proportionally to the photoionization heating, then the total heat input in the emitting gas should be doubled. {\sl This problem therefore deals with the global energetics of the early universe}. After reviewing previous models for \IZ\,NW (Sect.~\ref{prev_IZ}), observations and new photoionization models are described in Sects.~\ref{obs} \& \ref{newmod}. Results presented in Sect.~\ref{res} are discussed in Sect.~\ref{disc}. Concluding remarks appear in Sect.~\ref{concl}. Models for other GEHIIRs are reviewed in Appendix~A. Concepts undelying the new photoionization models are stated in Appendices~B and C. | \label{concl} Owing to its small heavy element content, \IZ\ stands at the high-\Te\ boundary of photoionized nebulae. Where ionization and temperature are sufficiently high, the cooling is little dependent on conditions, except through the concentration of H$^0$, controlled by density. Therefore, in the photoionization model logics, {\sl \Te\ is then a density indicator}, in the same way it is an O/H indicator in usual \hii\ regions. It is for not having recognized implications of this new logics, that low-metallicity BCD models failed. In a photoionization model study of \IZ\,NW, SS99 employed a filling-factor description and concluded that \Te(\oiii) was fundamentally unaccountable. The {\sl vogue} for this simple description of the ionized gas distribution resides in its apparent success for usual \hii\ regions, a success falling in fact to the strong dependence of cooling on abundances. Universally adopted along past decades, this concept led all authors to conclude that photoionization by hot stars did not provide enough energy to low-Z GEHIIRs. This conclusion is in line with a movement of calling into question photoionization by stars as the overwhelmingly dominant source of heat and ionization in gaseous nebulae, a movement cristalizing on the `$t^2$ problem' (Esteban et al. 2002; Peimbert et al. 2004), since the presence of \Te\ fluctuations {\sl supposedly larger} than those reachable assuming photoionization by stars implies additional heating. A conclusion of the present study is that the gas distribution is no less critical than the radiation source in determining the line spectrum of \hii\ regions. Assuming pure photoionization by stars, the remarkable piece of information carried by the large \Te(\oiii) of \IZ\,NW is that the mean density of the \oiii\ emitting region is much less than \Ne(\sii), a low \Ne\ confirmed by line ratios \oiv\la25.9$\mu$/\heii\la4686 and \feiii\la4986/\feiii\la4658. \IZ\,NW models comprising a plausible SED and respecting geometrical constraints can closely match almost all observed lines from UV to IR, including the crucial \oiii\la4363 (\siv\la10.5$\mu$ is a factor 2 off, however). Thus, extra heating by, \eg, dissipation of mechanical energy in the photoionized gas of low-metallicity BCD galaxies like \IZ\ is {\sl not} required to solve the `\Te(\oiii) problem'. Moreover, since low-ionization fine-structure lines can be explained by soft X-rays, (hydrodynamical) heating is {\sl not even} required in warm \hi\ regions protected from ionization and heating by star radiation. As a final note, on close scrutiny, the solutions found here {\sl tend to be just marginally consistent} with observed $r$(\oiii). Given the claimed accuracy in the different fields of physics and astrophysics involved, postulating a mechanical source of heating is premature, whereas a {\sl 2--3\% upward correction} to the collision strength for transition O$^{2+}$($^3$P\,--\,$^1$S) at \Te~$\sim$~2$\times$10$^4$\,K is an alternative worth exploring by atomic physics. Yet, another possibility is a substantial increase of the distance to \IZ. Owing to accurate spectroscopy and peculiar conditions in \IZ, important astrophysical developments are at stakes in the 5\% uncertainty attached to \oiii\ collision strengths. If photoionized nebulae are shaped by shocks and other hydrodynamic effects, this does not imply that the emission-line intensities are detectably influenced by the thermal energy deposited by these processes. Unravelling this extra thermal energy by means of spectroscopic diagnostics and models is an exciting prospect, whose success depends on a recognition of all resources of the photoionization paradigm. Adopting the view that photoionization by radiation from young hot stars, including WR stars, is the only excitation source of nebular spectra in BCD galaxies, yet without undue simplifications, may well be a way to help progresses in the mysteries of stellar evolution, stellar atmosphere structure, stellar supercluster properties, giant \hii\ region structure and, {\sl at last}, possible extra sources of thermal energy in BCDs. | 7 | 10 | 0710.5082 |
0710 | 0710.4400_arXiv.txt | {} {Bright HNC 1--0 emission, rivalling that of HCN 1--0, has been found towards several Seyfert galaxies. This is unexpected since traditionally HNC is a tracer of cold (10 K) gas, and the molecular gas of luminous galaxies like Seyferts is thought to have bulk kinetic temperatures surpassing 50 K. There are four possible explanations for the bright HNC: (a) Large masses of hidden cold gas; (b) chemistry dominated by ion-neutral reactions; (c) chemistry dominated by X-ray radiation; and (d) HNC enhanced through mid-IR pumping. In this work we aim to distinguish the cause of the bright HNC and to model the physical conditions of the HNC and HCN emitting gas.} {We have used SEST, JCMT and IRAM 30m telescopes to observe HNC 3--2 and HCN 3--2 line emission in a selection of 5 HNC-luminous Seyfert galaxies. We estimate and discuss the excitation conditions of HCN and HNC in NGC~1068, NGC~3079, NGC~2623 and NGC~7469, based on the observed 3--2/1--0 line intensity ratios. We also observed CN 1--0 and 2--1 emission and discuss its role in photon and X-ray dominated regions.} {HNC 3--2 was detected in 3 galaxies (NGC~3079, NGC~1068 and NGC~2623). Not detected in NGC~7469. HCN 3--2 was detected in NGC~3079, NGC~1068 and NGC~1365, it was not detected in NGC~2623. The HCN 3--2/1--0 ratio is lower than 0.3 only in NGC~3079, whereas the HNC 3--2/1--0 ratio is larger than 0.3 only in NGC~2623. The HCN/HNC 1--0 and 3--2 line ratios are larger than unity in all the galaxies. The HCN/HNC 3--2 line ratio is lower than unity only in NGC~2623, which makes it comparable to galaxies like Arp~220, Mrk~231 and NGC~4418. } {We conclude that in three of the galaxies the HNC emissions emerge from gas of densities $n\lesssim10^5~\3cm$, where the chemistry is dominated by ion-neutral reactions. The line shapes observed in NGC~1365 and NGC~3079 show that these galaxies have no circumnuclear disk. In NGC~1068 the emission of HNC emerges from lower ($<10^5~\3cm$) density gas than HCN ($>10^5~\3cm$). Instead, we conclude that the emissions of HNC and HCN emerge from the same gas in NGC~3079. The observed HCN/HNC and CN/HCN line ratios favor a PDR scenario, rather than an XDR one, which is consistent with previous indications of a starburst component in the central regions of these galaxies. However, the $N({\rm HNC})/N({\rm HCN})$ column density ratios obtained for NGC~3079 can be found only in XDR environments.} | The hydrogene cyanide, HCN molecule, is commonly used as an extragalactic tracer of molecular gas with densities $n$(H$_2$) larger than $10^4~\3cm$ (e.g. Solomon et al. \cite{solomon92}; Curran et al. \cite{curran00}; Kohno \cite{kohno05}). The HCN to CO intensity ratio varies significantly, from 1/3 to 1/40 in starburst galaxies, and it has not been determined whether this variation depends on the dense molecular gas content or on the abundance and/or excitation conditions. In addition, recent results seems to indicate that HCN may not be an unbiased tracer of the dense molecular gas content in LIRGs and ULIRGs (Graci\'a-Carpio et al. \cite{gracia06}). It is essential, therefore, to use other molecular tracers than HCN, in order to understand the physical conditions in the dense gas. A molecule of particular interest, for comparison with HCN, is its isomer HNC. The detection of interstellar HNC supports the theory of dominant ion-molecule chemistry in dark molecular clouds. Both species are thought to be created by the same dissociative recombination of HCNH$^+$. This ion can produce HCN and HNC, with approximately equal abundances. Models based only on this scheme would predict then an HNC/HCN ratio $\approx 1$. However, the observed HNC/HCN abundance ratios vary significantly between different kinds of molecular clouds - the ratio ranges from 0.03 to 0.4 in warm cores ($T_k > 15$ K), and can be as high as 4.4 in cold cores ($T_k < 15$ K). The CN (cyanogen radical) molecule is another tracer of dense gas, with a lower (by a factor of 5) critical density than HCN. CN is also chemically linked to HCN and HNC by photodissociations (e.g. Hirota \& Yamamoto \cite{hirota99}). Surveys of the 1--0 transition of CN and HNC have been done in order to trace a cold, dense phase of the gas in luminous galaxies (Aalto et al. \cite{aalto02}). It was found that the HNC 1--0 luminosities often rivalled those of HCN 1--0. These results seem to contradict the idea of warm ($T_k \gtrsim 50$ K) gas in the centers of luminous galaxies (e.g. Wild et al. \cite{wild92}; Wall et al. \cite{wall93}) whose IR luminosities were suggested to originate from star formation rather than AGN activity (Solomon et al. \cite{solomon92}). \begin{table*}[!t] \begin{minipage}{18cm} \caption[]{Sample of galaxies$^{~\rm a}$.} \label{tab:galaxies} \centering \begin{tabular}{lccccccc} \hline \noalign{\smallskip} Galaxy & Seyfert & RA & DEC & v$_{sys}$ & Distance$^{~\rm b}$ & $\Omega_S$(CO)$^{~\rm c}$ & $\Omega_S$(HCN)$^{~\rm d}$\\ & & [hh mm ss] & [$^{\circ}~{}'~{}''$] & [\kms] & [Mpc] & [$''~^2$] & [$''~^2$]\\ \noalign{\smallskip} \hline \noalign{\smallskip} NGC~3079 & 2 & 10 01 57.805 & +55 40 47.20 & 1116$\pm$1 & 15.0$\pm$1.1 & $15\times 7.5$ & $5\times 5$ \\ NGC~1068 & 2 & 02 42 40.711 & -00 00 47.81 & 1137$\pm$3 & 15.3$\pm$1.1 & $30\times 30$ & $10\times 10$ \\ NGC~2623 & 2 & 08 38 24.090 & +25 45 16.80 & 5549$\pm$1 & 74.6$\pm$5.4 & $8\times 8$ & $2.6\times 2.6$ \\ NGC~1365 & 1.8 & 03 33 36.371 & -36 08 25.45 & 1636$\pm$1 & 22.0$\pm$1.6 & $50\times 50$ & $16.5\times 16.5$ \\ NGC~7469 & 1.2 & 23 03 15.623 & +08 52 26.39 & 4892$\pm$2 & 65.8$\pm$4.8 & $8\times 8$ & $4\times 6$ \\ \noalign{\smallskip} \hline \end{tabular} \begin{list}{}{} \item[${\mathrm{a}}$)] The Seyfert classification, positions (in equatorial J2000 coordinates) and heliocentric radial velocities were taken from NED. \item[${\mathrm{b}}$)] The distances were calculated using the Hubble constant (H$_0 \approx 74.37$ \kms~Mpc$^{-1}$) estimated by Ngeow and Kanbur (\cite{ngeow06}). \item[${\mathrm{c}}$)] The source sizes of the CO 1--0 transition line were estimated from the maps presented in (Koda et al. \cite{koda02}) for NGC~3079, (Schinnerer et al. \cite{schinnerer00}) for NGC~1068, (Bryant et al. \cite{bryant99}) for NGC~2623, (Sandqvist \cite{sandqvist99}) for NGC~1365, and (Papadopoulos \& Allen \cite{papadopoulos00}) for NGC~7469. The source sizes for the $J = 2-1$ transition line were assumed equal to that of the $J = 1-0$ line. For NGC~7469, the source size of the CO 2--1 emission estimated from the corresponding map presented by Davies et al. (\cite{davies04}) agrees well with the source size estimated for the CO 1--0 line. \item[${\mathrm{d}}$)] Source sizes of HCN 1--0 were estimated from the corresponding maps published in (Kohno et al. \cite{kohno00}) for NGC~3079, (Kohno et al. \cite{kohno01} and Helfer \& Blitz \cite{helfer95}) for NGC~1068, (Davies et al. \cite{davies04}) for NGC~7469. The source sizes of NGC~2623 and NGC~1365 were estimated using proportions found in NGC~1068 (read text in \S3.5). Because of their chemical link, the source sizes of the CN and HNC molecules were assumed the same as that of HCN. Due to the lack of high resolution maps, the source sizes corresponding to the emission of the higher transition lines were assumed equal to that of the $J = 1-0$ line. \end{list} \end{minipage} \end{table*} According to observations in the vicinity of the hot core of Orion KL, experimental data, chemical steady state and shock models, the HNC/HCN ratio decreases as the temperature and density increase (e.g. Schilke et al. \cite{schilke92}; Talbi et al. \cite{talbi96}; Tachikawa et al. \cite{tachikawa03}). If a bright HNC 1--0 transition line is nevertheless detected under these conditions, it could be due to the following possible explanations: (a) the presence of large masses of hidden cold gas and dust at high densities ($n > 10^5~\3cm$); (b) chemistry dominated by ion-molecule reactions with HCNH$^+$ at low density ($n \approx 10^4~\3cm$) in regions where the temperature dependence of the HNC abundance becomes weaker; (c) enhancement by mid-IR pumping, also in low density regions where the lines would not be collisionally excited; and (d) the influence of UV-rays in Photon Dominated Regions (PDRs) and/or X-rays in X-ray Dominated Regions (XDRs) at densities $n \gtrsim 10^4~\3cm$ and at total column densities $N_{\rm H} > 3\times 10^{21}~\2cm$ (Meijerink \& Spaans \cite{meijerink05}). In the case of CN, observations of its emission towards the Orion A molecular complex (Rodr\'iguez-Franco et al. \cite{rodriguez98}) suggest that this molecule is also enhanced in PDRs, but particularly in XDRs, where a CN/HCN abundance ratio larger than unity is expected (e.g., Lepp \& Dalgarno \cite{lepp96} and Meijerink, Spaans, Israel \cite{meijerink07}). We have observed low and high transition lines of the HCN, HNC and CN molecules in a group of Seyfert galaxies, which are supposed to host both sources of power, AGN and starburst activity, in their central region. Our interest is to assess the excitation conditions of HCN and HNC, distinguish between the above possible causes of the bright HNC, and to explore the relation between the CN emission, XDRs and dense PDRs in these sources. In \S2 we describe the observations. The results (spectral lines, line intensities and line ratios) are presented in \S3. The interpretation of line shapes and gas distribution in the most relevant cases, as well as the possible explanations for the bright HNC and the modelling of the excitation conditions of HCN and HNC are discussed in \S4. The conclusions and final remarks of this work are presented in \S5. | We have used the SEST and JCMT telescopes to carry out a survey of CN 2--1, HCN 3--2 and HNC 3--2 line emission in a sample of 4 Seyfert galaxies, plus NGC~3079 which was observed with the IRAM 30m telescope. The conclusions we draw are as follows: \begin{list}{}{} \item[1)] We detected HNC 3--2 emission in 3 of the 5 galaxies, while we obtain an upper limit for one of them (NGC~7469). HCN 3--2 was also detected in 3 galaxies (NGC~3079, NGC~1068 and NGC~1365), while it was not detected in NGC~2623. CN 2--1, along with the spingroups (\textit{J} = 5/2 -- 3/2, \textit{F} = 7/2 -- 5/2) and (\textit{J} = 3/2 -- 1/2, \textit{F} = 5/2 -- 3/2) was also detected in NGC~3079, NGC~1068 and NGC~1365. \item[2)] {The line shapes} observed in NGC~1365 and NGC~3079 suggests that there is no circumnuclear disk in these galaxies. \item[3)] We find that in 3 of the galaxies the HNC 3--2/1--0 line ratios suggest that the HNC emissions emerge from gas of densities $n\lesssim10^5~\3cm$, where the chemistry is dominated by ion-neutral reactions. In NGC~2623 a model of large masses of hidden cold (10 K) gas and dust, as well as a chemistry dominated by ion-neutral reactions, are yet to be distinguished as the correct interpretation for the bright HNC observed in this galaxy. \item[4)] The 3--2/1--0 line ratios and the modelled excitation conditions imply that the HNC emission emerges from a more diffuse ($n<10^5~\3cm$) gas region than the HCN emission ($n>10^5~\3cm$) in NGC~1068, whereas they emerge from the same lower density ($n\lesssim10^5~\3cm$) gas in NGC~3079. \item[5)] The HCN/HNC and CN/HCN line ratios tentatively favor a PDR scenario, rather than an XDR one, in the 3 Seyfert galaxies where we have CN, HNC and HCN data. The $N({\rm HNC})/N({\rm HCN})$ column density ratios obtained for NGC~3079 can be found only in XDR environments. \end{list} In order to complete the sample, we plan to observe HCN 3--2 and CN 2--1 in NGC~7469, CN 2--1 in NGC~2623 and HNC 3--2 in NGC~1365. We plan to perform high resolution observations to further study the distribution and source sizes of CN and HNC. Modeling of the collision data for the CN molecule would be useful to estimate the $N({\rm CN})/N({\rm HCN})$ column density ratio, which would complement the $N({\rm HNC})/N({\rm HCN})$ ratio in order to have a more sophisticated tool to estimate and distinguish the prevalent environment conditions of the high density gas in the nuclear region of Seyfert galaxies. The AGN contribution (through XDR effects) is typically of a small angular scale and can be seriously affected by beam dilution at the transition lines studied in this work. On the other hand, the starburst contribution is of a larger angular scale than the AGN, and it effects can be contaminating our observations, and hence leading to the favored PDR scenario found with our models. Hence, our suggested interpretations could change if we zoom in on these sources. Therefore, high resolution maps of HNC and CN molecules are necessary to complement those of HCN, and to do a more accurate estimate of molecular abundances and line intensity ratios, which take source size into account. Observations of the higher transition lines (e.g. $J$=4--3) can also aid to disentangle the effects of the AGN and the starburst ring, due to the smaller beam size obtained at higher frequencies. | 7 | 10 | 0710.4400 |
0710 | 0710.3685_arXiv.txt | { Relic neutralinos produced after the Big Bang are favoured candidates for Dark Matter. They can accumulate at the centre of massive celestial objects like our Sun. Their annihilation can result in a high-energy neutrino flux that could be detectable as a localised emission with earth-based neutrino telescopes like ANTARES. In this paper a brief overview of the prospects of the indirect search for Dark Matter particles with the ANTARES detector will be given. The analysis method and expected performance for the detection of the expected neutrinos will be discussed. \PACS{ {95.35.+d}{Dark matter (stellar, interstellar, galactic, and cosmological)} \and {95.55.Vj}{Neutrino, muon, pion, and other elementary particle detectors; cosmic ray detectors} } % } % | \label{intro} It is now a well-established fact that according to our present theory of gravity, 85\%~of the matter content of our universe is missing. Observational evidence for this discrepancy ranges from macroscopic to microscopic scales, e.g. gravitational lensing in galaxy clusters, galactic rotation curves and fluctuations measured in the Cosmic Microwave Background. This has resulted in the hypothesised existence of a new type of matter called Dark Matter. Particle physics provides a well-motivated explanation for this hypothesis: The existence of (until now unobserved) massive weakly interacting particles (WIMPs). A favorite amongst the several WIMP candidates is the neutralino, the lightest particle predicted by Supersymmetry, itself a well-motivated extension of the Standard Model. If Supersymmetry is indeed realised in Nature, Supersymmetric particles would have been copiously produced at the start of our Universe in the Big Bang. Initially these particles would have been in thermal equilibrium. After the temperature of the Universe dropped below the neutralino mass, the neutralino number density would have decreased exponentially. Eventually the expansion rate of the Universe would have overcome the neutralino annihilation rate, resulting in a neutralino density in our Universe today similar to the cosmic microwave background. These relic neutralinos could then accumulate in massive celestial bodies in our Universe like our Sun through weak interactions with normal matter and gravity. Over time the neutralino density in the core of the object would increase considerably, thereby increasing the local neutralino annihilation probability. In the annihilation process new particles would be created, amongst which neutrinos. This neutrino flux could be detectable as a localised emission with earth-based neutrino telescopes like ANTARES. This paper gives a brief overview of the prospects for the detection of neutrinos originating from neutralino annihilation in the Sun with the ANTARES neutrino telescope. \begin{figure}[b] \center{ \includegraphics[width=0.45\textwidth,angle=0]{NEA_60kHz0XOFF_off.png} \caption{The ANTARES Neutrino Effective Area vs. $E_\nu$.} \label{fig:1} % } \end{figure} \begin{figure*}[t] \center{ \includegraphics[width=0.8\textwidth,angle=0]{psflux.png} \caption{Predicted $\nu_\mu+\bar{\nu}_\mu$ flux from the Sun in mSUGRA parameter space.} \label{fig:2} % } \end{figure*} | Nearly half of the ANTARES detector has been operational since January this year. The detector is foreseen to be completed in early 2008. The data show that the detector is working within the design specifications. As can be seen from Fig.~\ref{fig:4}, mSUGRA models that are excludable by ANTARES at 90\%~CL are found in the Focus Point region. The same models should also be excludable by future direct detection experiments, as is shown in Fig.~\ref{fig:7}. To improve the ANTARES sensitivity, a directional trigger algorithm has recently been implemented in the data acquisition system. In this algorithm, the known position of the potential neutrino source is used to lower the trigger condition. This increases the trigger efficiency, resulting in a larger $A_{\rm eff}^{\nu}$. In Fig.~\ref{fig:8}, the $A_{\rm eff}^{\nu}$ at the trigger level for the standard- and the directional trigger algorithm are shown in black (``{\em trigger3D}'') and red (``{\em triggerMX}'') respectively. | 7 | 10 | 0710.3685 |
0710 | 0710.2360_arXiv.txt | { The similarity of the observed mass densities of baryons and cold dark matter may be a sign they have a related origin. The baryon-to-dark matter ratio can be understood in the MSSM with right-handed (RH) neutrinos if CDM is due to a d = 4 flat direction condensate of very weakly coupled RH sneutrino LSPs and the baryon asymmetry is generated by Affleck-Dine leptogenesis along a d = 4 $\left(H_{u}L\right)^2$ flat direction. Observable signatures of the model include CDM and baryon isocurvature perturbations and a distinctive long-lived NLSP phenomenology. \PACS{ {98.80.Cq}{Cosmology} } % } % | \label{intro} A striking feature of the observed Universe is the similar mass density in baryons and cold dark matter. (The 'Baryon-to-Dark Matter' (BDM) ratio.) From the WMAP three-year results for the $\Lambda$CDM model, $\Omega_{DM}/\Omega_{B} = 5.65 \pm 0.58$ \cite{wmap}. However, in most models the physics of baryogenesis and dark matter production are unrelated. So why is the mass density in baryons within an order of magnitude of that of dark matter? There are three possibilities: \newline (i) A remarkable coincidence. \newline (ii) Some anthropic selection mechanism, usually assumed but undefined (e.g. in the case of thermal relic neutralino dark matter). \newline (iii) The mechanisms for the origin of the baryon asymmetry and dark matter are related. The latter possibility seems the simplest interpretation of the BDM ratio. Indeed, we may be ignoring a {\it big clue} to the nature of the correct BSM particle theory. It is highly non-trivial for a particle physics theory to have within its structure (without contrivance) a mechanism that can account for the BDM ratio. Therefore if the BDM ratio is due to such a mechanism it would provide us with a powerful principle by which to select the best canadiate particle physics models. BDM models broadly divide into two classes: \newline {\bf 1).Charge conservation based:} The dark matter particle and baryon number are related by a conserved charge, $Q_{B} + Q_{cdm} = 0 \Rightarrow n_{cdm} \sim n_{B}$. The CDM particle mass satisfies $ m_{cdm} = m_{n}n_{B}/n_{cdm}$ and so $m_{cdm} \sim 1 GeV$ is necessary. However, this does fit well with SUSY if the LSP mass is $O(m_{W})$ or larger. \newline {\bf 2). Dynamics based:} In this case the dark matter and baryon densities are related by {\it similar} physical mechanisms for their origin. This implies a less rigid relation between $n_{B}$ and $n_{CDM}$, which may allow us to understand why it is the mass rather than number densities that are observed to be similar. | RH sneutrino condensate CDM combined with Affleck-Dine baryogenesis can plausibly account for the observed similarity of the baryon and dark matter mass densities in the Universe. Seen as a selection principle, the requirement that a particle physics model can {\it without contrivance} account for the BDM ratio favours the MSSM with neutrino masses and RH sneutrino condensate CDM. It is quite remarkable that the MSSM with neutrino masses has the ability to account for the BDM ratio as a natural consequence its structure. CDM and baryon isocurvature perturbations are possible, the ratio of which gives information on the nature of SUSY inflation. Long-lived NLSP phenomenology is also expected, which can be distinguished from gravitino and axino LSP phenomenology via trapped stau final states, and from thermal relic RH sneutrino LSP phenomenology once the parameters of the MSSM are known. There are a number of issues which remain to be addressed. One is the possibility that $M_{N} \neq 0$. In this case it is possible that the heavier generation condensate RH sneutrinos, which will have a lifetime longer than the age of the Universe so long as $\lambda_{\nu}$ is still small, would decay into the LSP RH sneutrino plus $e^{+}e^{-}$. This could produce a potentially observable diffuse $\gamma$-ray background. In addition, the phenomenology and cosmology of MSSM-LSPs in this model should be studied in detail. | 7 | 10 | 0710.2360 |
0710 | 0710.2683_arXiv.txt | We study the dynamics of a twisted tilted disc under the influence of an external radiation field. Assuming the effect of absorption and reemission/scattering is that a pressure is applied to the disc surface where the local optical depth is of order unity, we determine the response of the vertical structure and the influence it has on the possibility of instability to warping. We derive a pair of equations describing the evolution of a small tilt as a function of radius in the small amplitude regime that applies to both the diffusive and bending wave regimes. We also study the non linear vertical response of the disc numerically using an analogous one dimensional slab model. For global warps, we find that in order for the disc vertical structure to respond as a quasi uniform shift or tilt, as has been assumed in previous work, the product of the ratio of the external radiation momentum flux to the local disc mid plane pressure, where it is absorbed, with the disc aspect ratio should be significantly less than unity. Namely, this quantity should be of the order of or smaller than the ratio of the disc gas density corresponding to the layer intercepting radiation to the mid plane density, $\lambda \ll 1$. When this condition is not satisfied the disc surface tends to adjust so that the local normal becomes perpendicular to the radiation propagation direction. In this case dynamical quantities determined by the disc twist and warp tend to oscillate with a large characteristic period $T_{*}\sim \lambda^{-1}T_{K}$, where $T_{K}$ is some 'typical' orbital period of a gas element in the disc. The possibility of warping instability then becomes significantly reduced. In addition, when the vertical response is non uniform, the possible production of shocks may lead to an important dissipation mechanism. | \noindent A significant number of X-ray binaries exhibit long-term periodicities on time-scales of ~10-100 d. Examples are Her X-1, SS 433 and LMC X-4 see e.g. Clarkson et al. 2003. Precession and warping of a tilted accretion disc has been proposed as an explanation (Katz 1973; Petterson 1975) which has subsequently been found to have observational support (Clarkson et al. 2003). The effect of radiation pressure on a twisted tilted accretion disc was first considered by Petterson 1977a,b who noted that when radiation from the central source is absorbed at the disc surface and re-emitted, a potentially important torque will result. Iping \& Petterson 1990 later suggested that such torques determined the shape of the disc and its precession rate. Pringle 1996 subsequently showed that an initially axisymmetric thin disc could be unstable to warping as a result of interaction with an external radiation field (see e.g. Maloney et a. 1998; Ogilvie \& Dubus 2001 and Foulkes et al. 2006 for later developments). The analysis considered the disc to behave as a collection of rings, which could interact by transferring angular momentum through the action of viscosity, but otherwise behaved as if they were rigid. In particular, the vertical displacement or tilt was assumed to be essentially uniform and independent of the vertical coordinate. Thus the pressure applied at the surface at optical depth unity is assumed to be effectively communicated through the disc vertical structure so as to give a near uniform response. In this paper we extend the theoretical treatment of the interaction of a disc with an external radiation field to take account of possible significant departures of the tilt or displacement from uniformity in the vertical direction. One of our objectives is to determine the conditions under which the assumption of a uniform response is valid, and then to estimate some of the consequences when they are not satisfied, including potential additional dissipation resulting from non linear effects such as the production of shock waves. The latter is done using a one dimensional slab analogue model with the required high resolution in the vertical direction. Although we focus on the effects of a surface pressure induced by an external radiation field, very similar considerations are likely to apply when warps are induced by a surface pressure resulting from interaction with an external wind (e.g. Quillen, 2001) or surface forces produced by the interaction of the disc with an external magnetic field originating in the central star ( e.g. Pfeiffer \& Lai, 2004). We begin by considering the relevant issues using simple physical arguments. Following Pringle 1996 we consider a thin disc immersed in an external radiation field with mid plane initially coinciding with a Cartesian $(x,y)$ plane. It is supposed that the upper surface is parallel to the external rays so that there is initially no interaction with the radiation. The upper surface is then given a vertical elevation $h(r,\phi),$ where we now use polar coordinates $(r,\phi).$ As a result, a disc element with surface area $d{\cal A}$ absorbs momentum from the radiation field at a rate \be {\dot {\cal F}} = F_0 d{\cal A} \left(r{\partial (h/r)\over \partial r}\right).\ee Here $ F_0$ is the momentum flux at radius $r$ associated with the external radiation field. The factor in brackets gives the angle through which the local normal is rotated as a result of the perturbation (see Appendix \ref{A} for more details). Assuming the absorbed momentum is reradiated isotropically above the disc by the surface layers at an optical depth of unity, there will be an applied pressure there of magnitude \be {\cal P} = {2F_0\over 3} \left(r{\partial (h/r)\over \partial r}\right).\ee This external pressure, when applied to a complete elementary ring, produces a net torque per unit length of magnitude \be { d {\cal T} \over dr} = {2\pi F_0 \over 3} \left(r^2{\partial (h/r)\over \partial r}\right){\bf l},\ee where the complex vector ${\bf l}$ has components equal to $(i,1,0)$ in the Cartesian coordinate system. Here, in performing the azimuthal integration, we take into account that the azimuthal dependence of $h$ is through a factor of the form $ \exp(-i\phi)$ and work with the radial amplitude from now on. To find the consequent evolution of the disc, one requires the component of the angular momentum per unit length perpendicular to its unperturbed direction. This is given by \be {d{\cal J}/dr} = 2\pi\Sigma r^2 \Omega {\langle h \rangle}i{\bf l},\ee with $\Omega$ and $\Sigma $ being the near Keplerian local disc angular velocity and the disc surface density, respectively. $\langle h \rangle =\int d\zeta \rho h /\Sigma $ where $\zeta $ is a vertical coordinate and $\rho $ is the gas density. For an elementary ring the condition that the rate of change of angular momentum equal the applied torque gives a tilt evolution equation of the form \be {\partial {\langle h \rangle}\over \partial t}= -i{ F_0 \over 3\Omega\Sigma} \left({\partial (h/r)\over \partial r}\right). \ee Assuming the vertical response is uniform, so that we can set $h= {\langle h \rangle} \equiv r{\bf W}$, we obtain a description of warp evolution equivalent to Pringle 1996 when effects due to viscosity and bending wave propagation are neglected (see sections \ref{simpa} and \ref{Dyneq} below). This description indicates the possibility of instability to radiation pressure warping. However, here we stress the fact that the averaged elevation $\langle h \rangle ,$ enclosed in angled brackets, applies to the bulk of the inertia of the disc and can therefore be shown to be close to the elevation of the mid plane. On the other hand $h$ as used in the torque formula applies to the disc surface elevation. An important aspect of this paper is to distinguish these two elevations and investigate under what conditions they can be taken to be equal. This would be possible if the vertical structure responds as a rigid body to the external pressure forcing and we find the conditions for this to occur. The general requirement is found to be that the density at the surface of the disc where the pressure is applied should not be too small. It is possible to estimate in a simple way when the vertical displacement response of the disc to the external pressure forcing becomes non uniform. Let us suppose that the external pressure $ {\cal P}$ is applied at the surface layer where the density $\rho = \rho_*.$ The induced vertical displacement $h,$ now should be considered to be also a function of the vertical coordinate $\zeta.$ Assuming a linear response, which should be appropriate for sufficiently small elevations and, for simplicity an isothermal equation of state with sound speed $c_s,$ it can be easily shown that in the upper layers of the disc where vertical motions dominate the Lagrangian pressure perturbation has the form determined by presence of $h$, $\Delta P = -\rho_{*}c_{s}^{2} {\partial h\over \partial \zeta}$, see equation (\ref{e24}) below. Equating this to the external pressure we have \be {\partial h\over \partial \zeta}= -{{\cal P}\over \rho_* c_s^2}.\ee If the vertical extent of the disc is $\zeta_0,$ and $h$ varies on a radial scale comparable to $r,$ the characteristic change in $h$ induced over the vertical thickness is easily estimated to be \be \Delta h \sim \zeta_0{{\cal P}\over \rho_* c_s^2} \sim {2F_0r \over 3\rho_* \Omega^2\zeta_0} \left({\partial (h/r)\over \partial r}\right),\label{i1}\ee where we use the approximate relation $c_s=\Omega\zeta_0.$ From the condition $\Delta h \sim h $, using (\ref{i1}) and estimating ${\partial h\over \partial r}\sim h/r$, it is clear that whether $h$ is uniform or not is governed by the the parameter $ F_0 r /(\rho_* \Omega^2\zeta_0^3)(\zeta_0/r)^2 = \lambda^{-1} \epsilon \delta^2,$ with $\lambda =\rho_*/\rho_c$ and $\delta = \zeta_0/r$ so defining $\epsilon= F_0 r /(\rho_c \Omega^2\zeta_0^3).$ For a uniform response, one requires that $\epsilon \delta ^2 \ll \lambda.$ This is equivalent to the requirement that the product of the ratio of the external radiation momentum flux to the local disc pressure with the disc aspect ratio should be significantly less than unity. We recall here that because of the geometrical configuration, the momentum flux locally reradiated by the disc is in general much less than the external momentum flux at that location. In this paper we extend the treatment of disc warping induced by the action of an external pressure to include the effects of the response of the vertical structure particularly when this is non uniform as is the case when the above condition is not satisfied. In that case we find that the disc dynamics enters a different regime. This is such that the exposed surface tends to align so as to reduce the momentum absorption and hence the applied pressure. In the extreme limit of this regime the upper surface acts as if it is in contact with a rigid wall with the tendency to radiation warping instability tending to vanish. In this limit quantities characterising the disc twist and warp tend to oscillate at typical frequency $\sim \lambda \Omega $. To illustrate these effects we derive a description of the one dimensional evolution of the disc inclination in radius and time that incorporates the effects of the vertical structure response, radiation torques and which applies both to the high viscosity regime, when the evolution is diffusive, and to the low viscosity regime when the evolution is wavelike. In all cases, the efficacy of surface radiation pressure driven instabilities is found to be reduced once the model parameters are such that the vertical response is significantly non uniform. We go on to estimate conditions for the response to be nonlinear and investigate the development of shock waves in the response using a one dimensional slab analogue model which has the same characteristic behaviour of linear perturbations as the full disc model. We find that such shocks potentially provide an important dissipation mechanism. The plan of this paper is as follows. In section \ref{sec1} we describe some aspects of the thin disc model used. In section \ref{Coords} we go on to introduce the twisted coordinate system used together with the notation convention, giving the basic equations in section \ref {Beq}. By integrating over the vertical direction we use these to derive a single equation governing the dynamics of a twisted disc in section \ref{Derveq}. When the disc behaves like a set of rigid rings for which the vertical response is uniform, this equation can be used to give a complete specification of the warp evolution. We note that this provides an extension of previous formulations to be able to consider the case when warps propagate as waves rather than diffuse radially (e.g. Nelson \& Papaloizou 1999). In this Paper we use the twisted coordinate system formalism first introduced by Petterson 1977a, 1978 where the dynamical equations take the most simple form. When this formalism is adopted and an accurate description of all components of the equations of motion is needed as in the problem on hand we show that, in general, there is an ambiguity in choice of twisted coordinates with a set of these describing the same physical situation. As discussed in section \ref{Derveq} a choice of a most appropriate twisted coordinate system can be motivated by the condition that perturbations of all dynamical quantities determined by the disc twist and warp are small. The transformation law between different twisted coordinates corresponding to the same physical situation is derived in appendix C for an inviscid disc. In order to obtain a complete description of the warp evolution when the vertical response is non uniform, we begin by obtaining a complete solution of the vertical problem for a polytropic model in section \ref{poly}. This is used to obtain a pair of equations governing the radial warp evolution. We also indicate how the results can be extended to apply to more general models. We confirm the condition for the disc response to be like that of a series of rigid rings as $\lambda \gg \epsilon (\zeta_0/r)^2.$ We go on to perform a linear stability analysis of the radial evolution equations adopting a WKB approach in section \ref {qualan} obtaining a maximum potential growth/decay rate in section \ref{7.3}. In section \ref{8} we discuss and confirm the analogy between the response of the disc and the linear and non linear dynamics of a vertically stratified one dimensional slab. We consider the development of shock waves and the formation of a rarefied hot atmosphere in section \ref{8.3} giving a crude estimate of the warp dissipation rate in section \ref{8.4}. Finally in section \ref {Conc} we discuss and summarize our results. | \label{Conc} In this paper we have presented a self-consistent calculation of the response of a twisted disc to the action of a radiation pressure force acting in upper layers of the disc. This is assumed to be due to the interception and subsequent reradiation of radiation from a central source. The radiation pressure force is assumed to be applied at a single density level $\rho_{*}$ corresponding to optical thickness unity. The degree of twist and warping is assumed to be small enough that linear theory can be used to calculate the response. The analysis of Pringle 1996 modelled the disc as a set of concentric rigid rings in a state of Keplerian rotation. These were assumed to communicate with each other through the exchange angular momentum because of the action of viscous forces. Up to now there has been no consideration of effects arising from the response of the disc vertical structure to the externally applied pressure. In this paper we have considered the warping and twisting of an accretion disc taking account the response of the disc vertical structure to an external radiation field assuming that the effect of self-shielding of radiation by the global twisted disc is not significant. We developed a description of the evolution of the disc in terms of a pair of equations governing the one dimensional evolution of the inclination as a function of radius and time (see section \ref{sec6} and equations (\ref{e26}), (\ref{e48}) and (\ref{e49})). This description extends earlier ones, so that the influence of radiation pressure on discs for which warps occur in the low viscosity bending wave regime, as well as the higher viscosity diffusive regime, may be considered. We found several qualitatively different dynamical regimes that may be realised in astrophysical discs. These are related to whether the character of the response of the disc vertical structure causes significant departures from what would be obtained for a rigid body. \subsection{ Conditions for a quasi-rigid response} We found that to avoid such departures, the momentum flux due to radiation from the central source $F_0,$ should be smaller than a critical value, $F_{*},$ given by $F_{*}=(\lambda /\delta )P_{c},$ where $\delta = \zeta_0/r $ is the disc aspect ratio, $P_{c}$ is the disc mid plane pressure and it has been assumed that the the ratio $(\lambda /\delta )$ is small (see sections \ref{Dyneq} and \ref{7.1.2}). Then, when $F_0 > F_{crit}\approx 0.1(\delta /\alpha )P_{c},$ warping instability of the disc becomes a possibility ( eg. Pringle, 1996). Thus for radiation driven warping of a disc that behaves like an ensemble of rigid rings, we find two requirements, namely that $F_{*} > F_{crit},$ together with $F_0 < F _{*}.$ These conditions together imply that the ration of the surface to mid plane density should be sufficiently large and such that $\lambda > \lambda_{crit}\approx 0.1\delta^{2}/\alpha.$ \subsection{ Large external radiation momentum flux} In the opposite limit of large $F_0 > max(F_{*}, F_{crit}),$ consideration of the disc structure in the vertical direction is essential as the vertical displacement ceases to be uniform. The perturbed gas motion in the disc is mainly determined by the vertical component of velocity and the perturbed quantities tend to oscillate at a characteristic frequency $\omega = (\rho_c \zeta_0 / 2\Sigma)\lambda \Omega $, with $\Omega$ being the local Keplerian angular velocity. In this situation, vertical motion tends to be suppressed by the external pressure field on the irradiated side of the disc while warping motions persist in the bulk of the disc and on the shielded side of the disc (see sections \ref{Dyneq} and \ref{smalllam}). The disc surface intercepting the radiation tends to align parallel to the radiation rays, thus decreasing the amount of intercepted radiation, while the inclination angles associated with the disc mid plane and the opposite free boundary of the disc oscillate, being approximately equal to each other. In this limit the upper surface of optical thickness one plays the role of an impenetrable wall. Thus the development of instability of the inclination angle due to radiation pressure effects (e.g. Pringle 1996) would be inhibited in this limit. In principle, the presence of warping motions in this regime would be associated with variability on a large characteristic time scale $T_{ch}\sim \lambda^{-1}T_{K}$, where $T_{K}$ and $\lambda $ are some 'typical' values of the orbital period and the density ratio. \subsection{Limits on the WKB growth rate} Following Pringle 1996 we have considered possible instabilities using a WKB approach. As this neglects potentially important global effects and boundary conditions results are not definitive, nonetheless they should give a good indication of when instability could be possible. We find that when $F_{0} < F_{*}$ the growth rate increases with $F_{0}$ while in the formal limit $F_{0} \rightarrow \infty$ it approaches zero. Thus in the linear theory there is an upper limit for the growth rate of the instability of the inclination angle which is always $\le (\rho_c \zeta_0/ \Sigma)\lambda\Omega $ in the absence of viscosity (see section \ref{7.3}). When present, viscosity acts to damp any instability at a rate $0.1\delta^2\Omega /\alpha,$ leading again to the condition $\lambda > \lambda_{crit}\approx 0.1\delta^{2}/\alpha$ for radiation warping instability to be feasible. \subsection{Conditions for non-linearity} Our results described above rely upon the applicability of linear perturbation theory. As discussed in section \ref{7.5}, the breakdown of linear theory is expected when the Lagrangian change of pressure induced in the surface layers becomes of the order of the unperturbed pressure even though the change of inclination angle could be very small. The corresponding constraints on the inclination angle are especially strong for disc models with a significant drop of temperature toward the boundaries of the disc. \subsection{Non linear calculations for the one dimensional slab analogue} A direct numerical approach to the issues discussed above is not yet feasible due to three-dimensional nature of the problem and the many physical processes involved. However, when vertical motions dominate, the problem becomes analogous to the one dimensional problem of vertical motions induced in a stratified gas column orbiting around a central source with Keplerian angular velocity. As discussed in Section \ref{8} the one dimensional slab gives the same fundamental period of oscillation as obtained from the full disc theory when an appropriate boundary condition on the upper surface of the column is specified. The dependence of the eigenfrequency on theoretical parameters as well as the effective presence, in the limit of small $\lambda,$ of an impenetrable wall at the upper surface of the column have been checked numerically. Where they can be compared, agreement between our analytical and numerical results is very good. We also used the one dimensional model in order to investigate possible consequences of non-linear behaviour in the system. To do this we studied the motion of the slab ensuing from the imposition of a vertical velocity profile with varying amplitude. We found that when the ratio of the Lagrangian pressure perturbation to the local value of the pressure at the upper boundary of the disc, $\Delta P/P_{*}$, becomes larger than $1-10,$ strong shocks propagating downward into the column are observed (see section \ref{8.3}). In principle, these shocks may lead to non-linear dissipation of energy stored in the vertical motion and in section \ref{8.4} we made a very crude estimate of a possible time scale of $200$ orbits for $\Delta P/P_{*}\sim 10$. However, this result must be checked in framework of a more sophisticated numerical approach which takes into account physical processes neglected in this study, most notably the effects of radiation transfer in the upper layers of the column. \subsection{Issues for future consideration} In a fully self-consistent study the dynamical effects of radiation pressure must be studied together with effects determined by the radiation heating of the disc atmosphere. This can be done in the most convenient way within the framework of the one dimensional model discussed above. In a realistic disc model, where effects due to the flaring of the disc lead to the interception and scattering of radiation in the disc photosphere, when axisymmetric and unperturbed, radiation heating can significantly influence on or even determine the value of the density ratio $\lambda.$ This parameter, being the ratio of the density at the absorbing surface to the mid plane density defines the boundary between different dynamical regimes for a twisted disc. In this connection it is interesting to note that in certain vertical models of X-ray heated accretion discs, the ratio $\lambda $ can be quite small, being of order of $10^{-4}-10^{-5}$, e.g. Jimenez-Garate et al 2002. In such models the new dynamical effects discussed in this Paper would certainly play an important role. | 7 | 10 | 0710.2683 |
0710 | 0710.0686_arXiv.txt | A survey of currently known extrasolar planets indicates that close to 20\% of their hosting stars are members of binary systems. While the majority of these binaries are wide (i.e., with separations between 250 and 6500 AU), the detection of Jovian-type planets in the three binaries of $\gamma$ Cephei (separation of 18.5 AU), GL 86 (separation of 21 AU), and HD 41004 (separation of 23 AU) have brought to the forefront questions on the formation of giant planets and the possibility of the existence of smaller bodies in moderately close binary star systems. This paper discusses the late stage of the formation of habitable planets in binary systems that host Jovian-type bodies, and reviews the effects of the binary companion on the formation of Earth-like planets in the system's habitable zone. The results of a large survey of the parameter-space of binary-planetary systems in search of regions where habitable planets can form and have long-term stable orbits are also presented. | 7 | 10 | 0710.0686 |
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0710 | 0710.2226_arXiv.txt | Gamma Ray Bursts (GRBs) show evidence of different spectral shapes, light curves, duration, host galaxies and they explode within a wide redshift range. However, the most of them seems to follow very tight correlations among some observed quantities relating to their energetic. If true, these correlations have significant implications on burst physics, giving constraints on theoretical models. Moreover, several suggestions have been made to use these correlations in order to calibrate GRBs as standard candles and to constrain the cosmological parameters. \\ We investigate the cosmological relation between low energy $\alpha$ index in GRBs prompt spectra and the redshift $z$. We present a statistical analysis of the relation between the total isotropic energy $E_{iso}$ and the peak energy $E_p$ (also known as Amati relation) in GRBs spectra searching for possible functional biases. \\ Possible implications on the $E_{iso}$ vs $E_p$ relation of the $\alpha$ vs $(1+z)$ correlation are evaluated. We used MonteCarlo simulations and the boostrap method to evaluate how large are the effects of functional biases on the $E_{iso}$ vs $E_p$. We show that high values of the linear correlation coefficent, up to about 0.8, in the $E_{iso}$ vs $E_p$ relation are obtained for random generated samples of GRBs, confirming the relevance of functional biases. \\ Astrophysical consequences from $E_{iso}$ vs $E_p$ relation are then to be revised after a more accurate and possibly bias free analysis. | Gamma-ray Bursts (GRBs) are brief and intense flashes of high energy radiation emitted mostly in the $\gamma$-ray band. They are detected from wholly random directions in the sky at the rate of about once a day and typically last from a few milliseconds to several minutes. Within a few years, the BATSE experiment on board the NASA's Compton Gamma Ray Observatory satellite (Fishman et al. 1989) has recorded over 2700 GRB events with an isotropic distribution in the sky (Meegan et al. 1996). However, although BATSE was very sensitive to high-energy photons, it could not discern the direction of a burst to better than a few degrees uncertainty, too large to pinpoint the location of individual explosions. The real revolutionary step forward occurred in the 1997, thanks to the Italian-Dutch BeppoSAX satellite (Boella et al. 1997). This satellite was not as sensitive as BATSE to $\gamma$ rays, but its relative quick response of pointing system, coupled with good accuracy position information, permitted the first detection of an X-ray {\it afterglow}, the radiation emitted after the initial burst of $\gamma$-ray (Costa et al. 1997). This discovery of afterglows made redshift measurements possible and confirmed that GRBs lie at cosmological distance ($0.0085$ (Galama et al. 1999) $<z<$ $6.29$ (Kawai et al. 2006)). It is well known that in the past years several correlations have been discovered linking various energies characterizing GRBs. All of them involve $E_p$, the energy peak of time integrated spectral energy distribution (SED). The first relation found links the rest frame isotropic energy $E_{iso}$ with $E_p$ (Amati et al. 2002, AM02 hereinafter, also known as 'Amati relation', see also Amati, 2006, AM06 hereinafter, for an updated version). This correlation is seen by several authors as an useful method to standardizing GRB energetics. Subsequently, $E_p$ was found to be tightly correlated also with the collimation corrected energy $E_{\gamma}$ (Ghirlanda et al. 2004, GH04 hereinafter, also known as 'Ghirlanda relation'), and this relation is used to constrain cosmological parameters using GRBs as 'known' candles. The same relations can be expressed in terms of luminosity using spectra not time integrated: $L_{iso}\propto E_{p}^k$ (Yonetoku et al. 2004), $L_{\gamma}\propto E_{p}^k$ (Ghirlanda, Ghisellini \& Firmani 2006). However, this kind of relations seems to contradict some observational evidences, as the large variety of light curves, spectra, redshifts, durations and host galaxies, leading us to suppose that the nature of GRB explosions is not unique. Band et al. (2005) and Nakar et al. (2005) tested the consistency of a large sample of BATSE GRBs (with unknown redshift) with the Amati and Ghirlanda relations. Their results suggest that they may be artifacts of selection effects, inferring that about half (Nakar et al. 2005) or even $\sim90\%$ (Band et al. 2005) of the whole GRB population cannot satisfy the correlation for any value of the redshift. However, these conclusions have been questioned by several other authors (Ghirlanda et al. 2005, Bosnjak et al. 2005, Pizzichini et al. 2005), that found instead that the peak energy and the fluences of BATSE GRBs with unknown redshift are fully consistent with the $E_{iso}~vs~E_{p}$ and the $E_{\gamma}~vs~E_{p}$ relations. In AM02, and in AM06, $E_{iso}$ and $E_{p}$ are not evaluated directly by a fitting procedure but they are calculated using analytic relations that include the same parameters. These procedures could introduce functional biases in the spectral correlation quoted above. The influence of these functional biases has not been never considered in literature. In this work we present a statistical analysis of the $E_{iso}$ vs $E_p$ relation, based on Monte Carlo and bootstrap simulations, searching for possible intrinsic correlation terms. We previously study the cosmological relation between the redshift $z$ and the low energy spectral index $\alpha$ (see Sec.3), because a correlation between these spectral parameters can make more tight the $E_{iso}$ vs $E_p$ relation. In Sect. 4 we evaluate functional and intrinsic correlation terms in the $E_{iso}$ vs $E_p$ relation, in order to estimate if the method used to derive it is statistically biased. | In this paper a statistical analysis of the GRBs parameters correlations $\alpha$ vs $(1+z)$ and $E_{iso}$ vs $E_p$ is presented. We studied the $\alpha$ vs $(1+z)$ correlation observing that it is present not only in the AM02 GRB sample but also in those analysed by GH04. As stated by different authors (Lloyd et al. 2000, Band et al. 1993), there is no clear explanation for it, but we suggested that it could be a direct observational consequence of a selection effects: faint high redshift GRBs are under the detectability threshold. As a consequence of this, the decrease of high redshift GRBs introduce a correlation between $\alpha$ and $z$. This correlation can have important consequences when we attempt to use GRBs as standard candles and, moreover, it introduces a bias in searching relations between quantities builded using $\alpha$ and $z$. We investigated how the presence of functional correlation terms affects the $E_{iso}$ vs $E_p$ relation. As builded in AM02, $E_{iso}$ and $E_p$ depend on a common set of spectral parameters. We performed several Monte Carlo simulations , initially considering only the intrinsic correlation term at a fixed redshift. We found that the $E_{iso}$ vs $E_p$ relation is biased, with a not negligible probability to reproduce correlation coefficents higher than 0.5. When we introduced in the generation of simulated sample also the $\alpha$ vs $(1+z)$ correlation, the mean value of the correlation coefficent increase up to $0.64$. Higher values were obtained using Eq.8 and Eq.9 instead of Eq.4 and Eq.5. A boostrap method was also applied to the AM02 sample. The good agreement with the previous analysis proves that our results are independent on the simulation code. An already discussed (e.g. Band et al. 2005) relevant subject is the contribution from selection effects in making more tight the $E_{iso}$ vs $E_p$ relation with respect to our simulations. It is possible that instrumental sensivity limits would produce forbidden regions in the plane $E_{iso}, E_p$. As an example, a combination of the spectral parameters in the GRB rest frame could make a GRB with a very low luminosity and therefore below the detectability threshold. We also stress that the SED energy peak is an averaged quantity, because during a single GRB that shows more bumps in its lightcurve we expect that $E_p$ changes, making difficult the link of this variable to physical quantities.\\ In the literature appear several other tight correlations concerning the energetics and the spectral distributions of GRBs (e.g $E_{\gamma}~vs~E_p$ (GH04), $L~vs~E_p$ (Yonetoku et al. 2004)), some of them are used to constrain cosmological parameters, using GRBs as 'known' candles. All of them are based on the $E_{iso}$ vs $E_p$ relation and this implies that they are affected by functional biases. In the GH04 relation, for example, the collimation corrected energy is computed considering the jet opening angle, that is in turn derived from $E_{iso}$ and $z$. \\ In any case, we do not exclude that a physical relation between the $E_{iso}$ and $E_p$ can really exist, but it can be safely established only after removing all the biases. A good solution could be to use spectral laws explicitly written in terms of the quantities for which a correlation is searched (as made, for example, in Tramecere, Massaro \& Cavaliere (2007) for the spectrum of the near HBL object Mkn 421). A further step could also be to work with homogeneous samples of GRBs in terms of some relevant characteristics (e.g. time evolution, spectral shape, ...) to verify that the correlation is actually followed by them. | 7 | 10 | 0710.2226 |
0710 | 0710.0015_arXiv.txt | We present observations of $^{13}$CO, C$^{18}$O, HCO$^{+}$, H$^{13}$CO$^{+}$, DCO$^{+}$ and N$_{2}$H$^{+}$ line emission towards the Barnard 68 starless core. The line profiles are interpreted using a chemical network coupled with a radiative transfer code in order to reconstruct the radial velocity profile of the core. Our observations and modeling indicate the presence of complex radial motions, with the inward motions in the outer layers of the core but outward motions in the inner part, suggesting radial oscillations. The presence of such oscillation would imply that B68 is relatively old, typically one order of magnitude older than the age inferred from its chemical evolution and statistical core lifetimes. Our study demonstrates that chemistry can be used as a tool to constrain the radial velocity profiles of starless cores. | \label{sec:introduction} Starless cores likely represent the earliest stage of the formation of a star. In the standard view of star formation, cores form out an initially magnetically sub-critical cloud, and evolve in a quasi-static fashion through ambipolar diffusion \citep{Shu87,Mouschiovas91}. In the opposite picture, cores are dynamic objects that form by shocks in a turbulent flow, fragment and collapse (or disappear) in a few crossing times \citep{Padoan01,Ballesteros-Paredes2003}. Velocity measurements can provide important constraints on these two scenarios. For example, \cite{Tafalla98} examined the kinematics of L1544 and inferred infall velocities up to 0.1 km s$^{-1}$ \emph{a priori} incompatible with sub-critical ambipolar diffusion models. However, \cite{Ciolek00} argued that these velocities can be understood if the core is super-critical. Line of sight velocities in cores can be established from observations of self-absorbed line profiles. In an infalling core, if the excitation temperature increases towards the center (as one can expect in a centrally condensed core with a roughly constant temperature), self-absorbed lines are expected to be asymmetric, with a blue peak brighter than the red peak \cite[see][]{Evans99}. This technique has been used in the past to detect collapse motions in starless cores and protostars \cite[see][for a review]{Myers00}. One of the major difficulties of this technique is the large degree of chemical complexity within these objects. For example, numerous molecules are observed to deplete in the densest regions of starless cores \cite[e.g.][]{Bergin02a,Tafalla02}, thus hampering our ability to measure the velocity in the dense central regions of these objects. Nevertheless, with a detailed knowledge of the chemistry of these objects, one can choose appropriate molecular transitions to trace different part of the cores, and unveil their radial kinematic structure \citep{Bergin03b,vanderTak05}. In this paper, we use line observations of $^{13}$CO, C$^{18}$O, HCO$^{+}$, H$^{13}$CO$^{+}$, DCO$^{+}$ and N$_{2}$H$^{+}$ to reconstruct the velocity profile of the Barnard 68 core (hereinafter B68). With a well defined physical \citep{Alves01,Bergin06b} and chemical \citep{Bergin02a,Maret06,Maret07a} structure, this core is an ideal target for such a study. For this we compare the observed line profiles with the predictions of a chemical network coupled with a radiative transfer code. The paper is organized as follow. Observations are presented in \S \ref{sec:observations}. The analysis is detailed in \S \ref{sec:analysis-results}, and the results are discussed in \S \ref{sec:discussion}. | \label{sec:discussion} Our observations and modelling indicate that the outer part of the core is collapsing while the inner part is expanding. The velocities in both regions are relatively small (a few tens of meters per second) and are largely sub-sonic. The transition between the collapsing and expanding region occurs at a radius of about 8,000 AU. It is important to note that the fit to the observations is probably not unique. However, large changes in the transition and velocities between the regions are likely ruled out. As already mentioned, our model does not reproduce the asymmetries of the N$_{2}$H$^{+}$ (1-0) and DCO$^{+}$ line profiles, due to an underestimating of the opacity. There are several possible explanations for this discrepancy. First, our chemical model might not predict the correct abundance profile for these two species. For example, increasing the DCO$^{+}$ abundance at intermediate radii, with a subsequent reduction at small radii to keep the column density constant would increase the opacity of line without changing the line flux significantly (which is correctly predicted). Indeed, our model predicts a DCO$^{+}$ (2-1) line intensity at $10 < A_{v} < 20$ that is lower than observed \citep[see Fig. 3 in][]{Maret07a}. Second, the physical structure of the core is also somewhat uncertain. For example, the temperature profile may be slightly different than the one predicted by \cite{Bergin06b} \citep[see][]{Crapsi07}, although small changes ($\pm 3$ K) in the temperature profile were found to have little effect on the line intensities. Turbulence at the center of the core might also be different that what assumed here. The turbulent profile was determined by \cite{Bergin06b} using C$^{18}$O observations, which are not a good probe of the innermost regions of the core because of heavy depletion. Decreasing the turbulent velocity profile in the innermost region of the core would also increase the opacity of these two lines. Regardless of the discrepancy with our model predictions, we emphasize these two line profiles, together with the HCO$^{+}$ (4-3) line profile, unambiguously indicate the presence of outwards motions at the center of the core. We found that no constant infalling velocity profile could reproduce the observations. The step-like velocity profile we obtain is suggestive of the presence of radial oscillations in the core. Small non-radial oscillations of the outer layers of the core around a stable equilibrium have been proposed by \citet{Lada03} and \cite{Redman06} to interpret the asymmetry of the CS (2-1) and HCO$^{+}$ (3-2) lines observed in Barnard 68. \cite{Lada03} also suggested that, in addition to the surface oscillations seen in the CS (2-1) line, some radial oscillation might also be present. Our observations and modeling are in agreement with this hypothesis. \cite{Keto05} proposed a hydrodynamic model for the evolution of cores in quasi-equilibrium. In their modeling, a small perturbation of a stable core (according to the Bonnor-Ebert criterion) is found to result in damped oscillations around the equilibrium position. \cite{Keto06} showed that the surface pattern observed in the CS (2-1) line towards B68 is consistent with the (n = 1, l = m = 2) quadripolar oscillation mode for a specific orientation ($\lambda = \theta = 30^\circ$). For the same oscillation mode, the radial velocity is qualitatively similar to the one obtained in this study, but is out of phase: the inner part of the core has a negative radial velocity, while the outer part of the core has a positive velocity. At the other half of the oscillation cycle, the velocity would be reversed and in qualitative agreement with the profile we obtain (E. Keto, \emph{priv. comm.}) As pointed out by \cite{Keto06} and \cite{Redman06}, the presence of oscillations in B68 suggests that the core is relatively old, typically older than a few sound speed crossing times ($\sim 10^{6}$ yr). This is comparable to the ambipolar diffusion timescale at the center of the core and more than one order of magnitude higher than the free-fall time scale \citep[$10^{6}$ and $7 \times 10^{4}$ yr respectively;][]{Maret07a}. This is also about an order of magnitude higher than the age of the core as determined from the degree of CO depletion \citep[$10^{5}$yr;][]{Bergin02a,Bergin06b}\footnote{Strictly speaking, the age determined by this method is a lower limit of the real age of the cloud, because CO is assumed to be pre-existing, and the density is supposed to be constant \citep{Bergin06b}.}. For $t = 3 \times 10^{5}$ yr, \cite{Bergin06b} model predicts a C$^{18}$O (1-0) line intensity that is three time smaller than the observations, while at $t = 10^{6}$ yr, the greater CO depletion leads to an even greater mismatch between observations and model predictions. The N$_{2}$H$^{+}$, HCO$^{+}$ H$^{13}$CO$^{+}$, and DCO$^{+}$ would also be underestimated by the model because of the freeze-out of their parent of their precursors, N$_{2}$, CO and $^{13}$CO. Thus the ``chemical age'' of B68 inferred from the CO depletion appears to be hard to reconcile with the ``dynamical age'' implied by the presence of surface oscillations. It is also interesting to compare the sound crossing time in B68 to typical statistical starless core lifetimes \cite[see][for a review]{WardThompson07}. Using sub-millimeter continuum emission maps, \cite{Kirk05} obtained lifetimes of $1-3 \times 10^{5}$ years which are in good agreement with the age of the B68 determined from CO depletion, but again about an order of magnitude higher than the sound crossing time. The present study emphasizes the importance of understanding the chemistry to constrain the kinematics of starless using spectral line profiles. Because of the strong chemical gradients that exist in these cores -- as well as excitation effects -- different lines can be used to selectively trace different parts of the core. We have demonstrated that, using lines that are appropriately chosen, it is possible to reconstruct the core radial velocity profile. In B68, our observations and modelling suggest the presence of complex line-of-sight motions that are consistent radial oscillations. The presence of such oscillation indicates that some core are long-lived, with lifetimes about an order of magnitude higher than those derived from their chemical evolution or sub-millimeter surveys. So far, oscillations have been inferred in only one other core, FeSt 1-457 \citep{Aguti07}. Clearly, similar studies on a larger number of cores are needed to establish if such cores are common. | 7 | 10 | 0710.0015 |
0710 | 0710.3400_arXiv.txt | {} { We investigate whether the globular clusters (GCs) in the recently published sample of GCs in the Fornax cluster by Bergond and coworkers are indeed intra--cluster objects.} {We combine the catalogue of radial velocity measurements by Bergond et al.~with our CTIO MOSAIC photometry in the Washington system and analyse the relation of metal--poor and metal--rich GCs with their host galaxies. } {The metal--rich GCs appear to be kinematically associated with their respective host galaxies. The vast majority of the metal--poor GCs found in between the galaxies of the Fornax cluster have velocities which are consistent with them being members of the very extended NGC\,1399 GC system. { We find that when the sample is restricted to the most accurate velocity measurements, the GC velocity dispersion profile can be described with a mass model derived for the NGC\,1399 GC system within 80\,kpc. We identify one ``vagrant'' GC whose radial velocity suggests that it is not bound to any galaxy unless its orbit has a very large apogalactic distance.}} {} {} | The existence of intra--cluster globular clusters (ICGCs), i.e. globular clusters (GCs) which are not bound to individual galaxies but, rather, move freely through the potential well of a galaxy cluster as a whole, was proposed by White~(\cite{white87}) and West et al.~(\cite{west95}). In the nearby ($D=19\,\rm{Mpc}$) Fornax cluster of galaxies, ICGC candidates were identified as an excess population of GC candidates in the vicinity of dwarf galaxies (Bassino et al.~\cite{bassino03}), and as a region of enhanced number density of GC candidates in between the central dominant galaxy NGC\,1399 and its second--nearest (giant) neighbour, NGC\,1387 (Bassino et al.~\cite{b06a}, cf.~their Fig.~9). Tamura et al.~(\cite{tamura06}) performed a wide--field survey of the central Virgo cluster and detected an excess population of GC candidates far away from any major galaxy. The spectroscopic confirmation of these candidate ICGCs, however, is still pending. \par Recently, Bergond et al.~(\cite{bergond07}) presented the velocities of a sample of GCs in the Fornax cluster, with some objects having projected distances of more than 230\,kpc from NGC\,1399. These authors labelled a significant fraction of their objects as ICGCs. Yet, one has to bear in mind that, with the photometric study by Bassino et al.~(\cite{b06a}), it became clear that the very populous globular cluster system (GCS) of the central galaxy, NGC\,1399, has an extent of at least 250\,kpc, which is comparable to the core radius of the cluster ($R_{King}=0.7^{\circ}\simeq 230\,\rm{kpc}$, Ferguson \cite{ferguson89}). \par In this \emph{Letter}, we present Washington photometry for 116 of the 149 GCs of the Bergond et al.~kinematic sample and demonstrate that the photometric and dynamical properties of the ``ICGCs'' are rather consistent with those of GCs belonging to NGC\,1399. | The red (metal--rich) GCs seem to be kinematically associated with their respective host galaxy or with NGC\,1399. The population of ICGCs is predominantly blue, as expected from the shallower number density profile of the metal--poor NGC\,1399 GCs. For the blue GCs, it is harder to determine whether a given GC ``belongs'' to a minor Fornax member or if it is part of the ICGC/NGC\,1399 population. We conclude that the vast majority of the blue ICGCs in fact belongs to the very extended NGC\,1399 GCS. Out to at least 165\,kpc, the velocity dispersion profile of the GCs is consistent with the mass profile derived from the dynamics of the NGC\,1399 GCS within 80\,kpc.\par We propose to distinguish between intra--cluster GCs i.e.~GCs which are found at large distances from the galaxies in a cluster (i.e.~a geometrical classification) and \emph{vagrant} GCs, which are characterised by radial velocities suggesting that they are not bound to any galaxy in particular, but rather belong to the galaxy cluster as a whole. The example of Fornax illustrates the difficulties one faces when trying to identify vagrant clusters. The detection of stripped GCs is made difficult by the richness and extent of the NGC\,1399 GCS and its high velocity dispersion. For NGC\,1404, which has a velocity of about $500\,\rm{km\,s}^{-1}$ w.r.t.~NGC\,1399, but a (projected) distance of only $\sim\!50\,\rm{kpc}$, it will be extremely hard to distinguish between a superposition along the line--of--sight and stripping. NGC\,1387, at about 100\,kpc has a systemic velocity which is just $\sim150\,\rm{km\,s}^{-1}$ lower than that of NGC\,1399, making it hard to single out stripped GCs, although the photometry of Bassino et al.~(\cite{b06a}) suggests their presence. We suggest that a dynamical study of the NGC\,1399/NGC\,1404 and NGC\,1387 region might yield evidence for tidal structures. | 7 | 10 | 0710.3400 |
0710 | 0710.1405_arXiv.txt | We analyze \ion{Fe}{1} 630~nm observations of the quiet Sun at disk center taken with the spectropolarimeter of the Solar Optical Telescope aboard the {\em Hinode} satellite. A significant fraction of the scanned area, including granules, turns out to be covered by magnetic fields. We derive field strength and inclination probability density functions from a Milne-Eddington inversion of the observed Stokes profiles. They show that the internetwork consists of very inclined, hG fields. As expected, network areas exhibit a predominance of kG field concentrations. The high spatial resolution of {\em Hinode}'s spectropolarimetric measurements brings to an agreement the results obtained from the analysis of visible and near-infrared lines. | \label{sec:intro} Most of the studied aimed at determining the distribution of field strengths in the internetwork (IN) quiet Sun have used polarimetric measurements in the spectral regions around 630~nm and 1565~nm, but their results do not agree. The visible \ion{Fe}{1} lines at 630.2~nm indicate a predominance of kG fields (S\'anchez Almeida \& Lites 2000; Dom\'{\i}nguez Cerde\~na et al.\ 2003; Socas-Navarro \& Lites 2004), whereas the infrared lines at 1565~nm suggest hG fields (Lin 1995; Lin \& Rimele 1999; Khomenko et al.\ 2003; Mart\'{\i}nez Gonz\'alez et al.\ 2006a; Dom\'{\i}nguez Cerde\~na et al.\ 2006). The distribution of IN field inclinations has only been studied by Lites et al.\ (1996) and Khomenko et al.\ (2003). Here we analyze \ion{Fe}{1} 630~nm measurements of the quiet Sun taken by the spectropolarimeter aboard {\em Hinode} at the unprecedented spatial resolution of 0\farcs32. The observed Stokes spectra are inverted to determine the distribution of field strengths and inclinations in the observed region. Our results show that most of the IN fields are weak, opposite to what has been found from ground-based measurements of the same lines at 1\arcsec. | The high spatial resolution spectropolarimetric measurements of {\em Hinode} indicate that most IN fields are weak. This is in agreement with the picture derived from the more magnetically sensitive \ion{Fe}{1} lines at 1565~nm (Lin 1995; Collados 2001; Khomenko et al.\ 2003; Mart\'{\i}nez Gonz\'alez et al.\ 2006b) and from lines showing hyperfine structure such as \ion{Mn}{1} 553~nm (L\'opez Ariste et al.\ 2006) and \ion{Mn}{1} 1526.2~nm (Asensio Ramos et al.\ 2007). Keller et al.\ (1994) also found weak fields in the internetwork using the \ion{Fe}{1} 525.0~nm lines, although at a lower spatial resolution and without inclination information. Our results seem to confirm the mean IN field strength of $\sim$100~G derived by Trujillo Bueno et al.\ (2004) from a Hanle-effect interpretation of \ion{Sr}{1} 460.7~nm measurements. Interestingly, the slope of the field strength distribution in the IN is similar to that obtained from magneto-convection simulations of comparable mean flux density. The observed field inclinations, however, turn out to be significantly larger than those predicted by the simulations. The scenario of an IN filled by nearly horizontal hG fields is compatible with the large trasverse magnetic fluxes found in the IN by Lites et al.\ (2007a,b). We still do not know the origin of such ubiquitous horizontal IN fields, but Lites et al.\ (2007b) have suggested a number of plausible mechanisms. In summary, Milne-Eddington inversions of the \ion{Fe}{1} 630~nm lines observed by {\em Hinode} at 0\farcs32 reveal a predominance of hG fields in quiet Sun internetwork regions, contrary to what is obtained from the same lines at 1\arcsec\/. This is the first time that \ion{Fe}{1} 630~nm observations confirm the weak IN fields indicated by near-infrared measurements, which may definitely close the discrepancy between the results derived from both spectral regions. | 7 | 10 | 0710.1405 |
0710 | 0710.3636_arXiv.txt | $UBVRI$ photometry and medium resolution optical spectroscopy of peculiar Type Ia supernova SN 2005hk are presented and analysed, covering the pre-maximum phase to around 400 days after explosion. The supernova is found to be underluminous compared to "normal" Type Ia supernovae. The photometric and spectroscopic evolution of SN 2005hk is remarkably similar to the peculiar Type Ia event SN 2002cx. The expansion velocity of the supernova ejecta is found to be lower than normal Type Ia events. The spectra obtained $\gsim 200$ days since explosion do not show the presence of forbidden [\ion{Fe}{ii}], [\ion{Fe}{iii}] and [\ion{Co}{iii}] lines, but are dominated by narrow, permitted \ion{Fe}{ii}, NIR \ion{Ca}{ii} and \ion{Na}{i} lines with P-Cygni profiles. Thermonuclear explosion model with Chandrasekhar mass ejecta and a kinetic energy smaller ($\KE = 0.3 \times 10^{51} {\rm ergs}$) than that of canonical Type Ia supernovae is found to well explain the observed bolometric light curve. The mass of \Nifs\ synthesized in this explosion is $0.18 \Msun$. The early spectra are successfully modeled with this less energetic model with some modifications of the abundance distribution. The late spectrum is explained as a combination of a photospheric component and a nebular component. | \label{sec:introduction} An impressive homogeneity in the light curves and peak luminosities make Type Ia supernovae (SNe Ia) good candidates in the determination of the extragalactic distance scale. Though a majority of the observed SNe Ia belong to the "normal" type (Branch et al.\ 1993), a number of studies indicate significant photometric as well as spectroscopic differences. For example, studies of nearby supernovae by Li et al.\ (2001) indicate that 64\% SNe Ia are "normal", while 20\% are of the overluminous SN 1991T type and 16\% of the underluminous SN 1991bg type. The peak absolute luminosities of SNe Ia are well correlated with their immediate post-maximum decline rate, forming a photometric sequence from the luminous blue events with a relatively slow decline rate to the faster, red, subluminous events (Hamuy et al.\ 1996a,b; Phillips et al.\ 1999). However, there are a few SNe Ia that are known to deviate from this relation, the most notable amongst them being SN 2002cx. This supernova was found to be underluminous, but had a light curve decline rate $\Delta$m$_{15}(B)=1.29$, comparable to normal SNe (Li et al.\ 2003). The early phase ($\lsim 100$ days after explosion) spectra indicate line velocities lower by a factor of 2 compared to those of normal SNe Ia (Li et al.\ 2003; Branch et al.\ 2004). Furthermore, the late phase ($\sim250$ days after explosion) spectra are also quite dissimilar compared to normal SNe Ia and possibly consist of P-Cygni profiles (Jha et al.\ 2006). Interestingly, SN 2002cx is not a unique event. Jha et al.\ (2006) and Phillips et al.\ (2007) have shown that SN 2005hk shows photometric and spectroscopic behaviour almost identical to that of SN 2002cx. Further, Jha et al.\ (2006) have shown that SNe 2003gq and 2005P also show spectra similar to SN 2002cx. Spectropolarimetric observations of SN 2005hk at the early phases by Chornock et al.\ (2006) indicate low polarization levels, indicating that the peculiarities of SN 2002cx-like SNe do not result from an extreme asphericity. Based on the analyses of the early phase spectra of SN 2002cx, Branch et al.\ (2004) suggest that the observed lower line velocities are consistent with the deflagration models of explosion. Jha et al.\ (2006) report a possible detection of \ion{O}{i} lines in the late phase spectra and suggest large scale mixing in the central region, which is also consistent with three-dimensional (3D) deflagration models. Phillips et al.\ (2007) find that, qualitatively the observed light curves of SN 2005hk are in reasonable agreement with model calculations of a 3D deflagaration model that produces $\sim 0.2~\Msun$ of $^{56}$Ni. Understanding the nature of this class of SNe Ia is thus quite important for the study of homogeneity and heterogeneity of SNe Ia. It may provide a caution for the cosmological use of SNe Ia and a strong constraint to the explosion models. We present in this paper the photometric and spectroscopic development of SN 2005hk over $\sim400$ days since explosion. | Photometric and spectrosopic data on the peculiar SN 2005hk are presented. The $B$ band light curve of SN 2005hk shows comparable pre-maximum brightening, faster decline in the initial $\sim$20 days past maximum and slower decline beyond $\sim$50 days past maximum, as compared to normal Type Ia supernovae. The fainter peak bolometric luminosity indicates synthesis of small amount of \Nifs \ in the explosion. Low expansion velocity of the ejecta together with fainter peak luminosity is explained by an explosion with lower kinetic energy. A reasonable fit to the bolometric light curve of SN 2005hk is achieved with a less energetic ($0.30 \times 10^{51}$erg) model which synthesized $0.18 \Msun$ of \Nifs. The light curve evolution is similar to SN 2002cx. The pre-maximum spectrum of SN 2005hk is similar to that of SN 1991T, with much lower expansion velocity. The spectral evolution of SN 2005hk is very similar to SN 2002cx. The spectrum of SN 2005hk at late phases ($> 200$ days) is similar to SN 2002cx, except for higher expansion velocities and higher velocity dispersions. The presence of P-Cygni profiles in the late phase spectrum of SN 2005hk indicates that the ejecta have not become optically thin till our last observation. Modeling of the pre-maximum spectra of SN 2005hk indicates a relatively higher temperature which makes \ion{Fe}{iii} lines strong. The presence of weak \ion{O}{i} line at $\lambda$7774 at almost all epochs is modeled as a consequence of high abundance of completely mixed unburned oxygen in the ejecta. The late phase spectra of SN 2005hk are modeled as a combination of the photospheric and nebular components, with the nebular line emitting region being more centrally concentrated than is expected in the case of a lower energetic model. | 7 | 10 | 0710.3636 |
0710 | 0710.1069_arXiv.txt | We utilize existing imaging and spectroscopic data for the galaxy clusters MS2137-23 and Abell 383 to present improved measures of the distribution of dark and baryonic material in the clusters' central regions. Our method, based on the combination of gravitational lensing and dynamical data, is uniquely capable of separating the distribution of dark and baryonic components at scales below 100 kpc. Our mass models include pseudo-elliptical generalized NFW profiles for constraining the inner dark matter slope, and our lens modeling takes into account both the ellipticity and substructure associated with cluster galaxies as necessary in order to account for the detailed properties of multiply-imaged sources revealed in Hubble Space Telescope images. We find a variety of strong lensing models fit the available data, including some with dark matter profiles as steep as expected from recent simulations. However, when combined with stellar velocity dispersion data for the brightest member, shallower inner slopes than predicted by numerical simulations are preferred, in general agreement with our earlier work in these clusters. For Abell 383, the preferred shallow inner slopes are statistically a good fit only when the multiple image position uncertainties associated with our lens model are assumed to be 0\farcs5, to account for unknown substructure. No statistically satisfactory fit was obtained matching both the multiple image lensing data and the velocity dispersion profile of the brightest cluster galaxy in MS2137-23. This suggests that the mass model we are using, which comprises a pseudo-elliptical generalized NFW profile and a brightest cluster galaxy component may inadequately represent the inner cluster regions. This may plausibly arise due to halo triaxiality or by the gravitational interaction of baryons and dark matter in cluster cores. The intriguing results for Abell 383 and MS2137-23 emphasize the need for a larger sample of clusters with radial arcs. However, the progress made via this detailed study highlights the key role that complementary observations of lensed features and stellar dynamics offer in understanding the interaction between dark and baryonic matter on non-linear scales in the central regions of clusters. | Cold dark matter (CDM) simulations (both with and without the inclusion of baryonic physics) are a crucial tool and proving ground for understanding the physics of the universe on nonlinear scales. One of the most active aspects of research in this area concerns the form of the dark matter density profile. Key questions raised in recent years include: Is there a universal dark matter density profile that spans a wide range of halo masses? What is the form of this profile and how uniform is it from one halo to another? To what extent do baryons modify the dark matter distribution? Drawing on a suite of N-body simulations, \citet{NFW97} originally proposed that the dark matter density profiles in halos ranging in size from those hosting dwarf galaxies to those with galaxy clusters have a universal form. This 3-D density distribution, termed the NFW profile, follows $\rho_{DM}\propto r^{-1}$ within some scale radius, $r_{sc}$, and falls off as $\rho_{DM}\propto r^{-3}$ beyond. Subsequent simulations indicated that the inner density profile could be yet steeper - $\rho_{DM}\propto r^{-1.5}$ \citep{M98,Ghigna00}. As computing power increases and numerical techniques improve, it is now unclear whether the inner dark matter distribution converges to a power law form rather than becoming progressively shallower in slope at smaller radii \citep{P03,Navarro04,Diemand04,Diemand05}. For comparisons with data, such simulations need to account for the presence of baryons. This is particularly the case in the cores of rich clusters. Although baryons represent only a small fraction of the overall cluster mass, they may be crucially important on scales comparable to the extent of typical brightest cluster galaxies. Much work is being done to understand the likely interactions between baryons and DM \citep{Gnedin04,Nagai05,Faltenbacher05}. These simulations will provide refined predictions of the relative distributions of baryons and DM. This paper is a further step in a series which aims to present an observational analog to progress described above in the numerical simulations. At each stage it is desirable to confront numerical predictions with observations. Whereas some workers have made good progress in constraining the {\em total} density profile (e.g.~\citet{Broadhurst05b}), in order to address the relevance of the numerical simulations we consider it important to develop and test techniques capable of separating the distributions of dark and baryonic components (e.g. ~\citet{Sand02,Zappacosta06,Biviano06,Mahdavi07}). This paper presents a refined version of the method first proposed by \citet{Sand02}, exploited more fully in \citet{Sand04} (hereafter S04). S04 sought to combine constraints from the velocity dispersion profile of a central brightest cluster galaxy (BCG) with a strong gravitational lensing analysis in six carefully selected galaxy clusters in order to separate the baryonic and dark matter distributions. S04 carefully selected clusters to have simple, apparently 'relaxed' gravitational potentials in order to match broadly the 'equilibrium' status of the simulated dark matter halos originally analyzed by \citet{NFW97} and subsequent simulators. For example, Abell 383 and MS2137-23 have almost circular BCGs ($b/a$=0.90 and 0.83 respectively), require a single cluster dark matter halo to fit the strong lensing constraints (in contrast to the more typical clusters that require a multi-modal dark matter morphology -- Smith et al. 2005), have previously published lens models with a relatively round dark matter halo ($b/a$=0.88 and 0.78 respectively - Smith et al. 2001; Gavazzi 2005), and display no evidence for dynamical disturbance in the X-ray morphology of the clusters (Smith et al. 2005; Schmidt \& Allen 2006). The merit of the approach resides in combining two techniques that collectively probe scales from the inner $\sim$10 kpc (using the BCG kinematics) to the $\sim$100 kpc scales typical of strong lensing. Whereas three of the clusters contained tangential arcs, constraining the total enclosed mass within the Einstein radius, three contained both radial and tangential gravitational arcs, the former providing additional constraints on the derivative of the total mass profile. In their analysis, S04 found the gradient of the inner dark matter density distribution varied considerably from cluster to cluster, with a mean value substantially flatter than that predicted in the numerical simulations. S04 adopted a number of assumptions in their analysis whose effect on the derived mass profiles were discussed at the time. The most important of these included ignoring cluster substructure and adopting spherically-symmetric mass distributions centered on the BCG. The simplifying assumptions were considered sources of systematic uncertainties, of order 0.2 on the inner slope. Although the six clusters studied by S04 were carefully chosen to be smooth and round, several workers attributed the discrepancy between the final results and those of the simulations as likely to arise from these simplifying assumptions \citep{Bartelmann04,Dalal04b,Meneghetti05}. The goal of this paper is to refine the data analysis for two of the clusters (MS2137-23 and Abell 383) originally introduced by S04 using fully 2-D strong gravitational lensing models, thus avoiding any assumptions about substructure or spherical symmetry. The lensing models are based on an improved version of the LENSTOOL program (\citealt{Kneibphd,Kneib96}; see Appendix; http://www.oamp.fr/cosmology/lenstool/). A major development is the implementation, in the code, of a pseudo-elliptical parameterization for the NFW mass profile, i.e. a generalization of those seen in CDM simulations, viz: \begin{equation}\label{eqn:gnfw} \label{eq:gnfw} \rho_d(r)=\frac{\rho_{c} \delta_{c}}{(r/r_{sc})^{\beta}(1+(r/r_{sc}))^{3-\beta}} \end{equation} where the asymptotic DM inner slope is $\beta$. This formalism allows us to overcome an important limitation of previous work and takes into account the ellipticity of the DM halo and the presence of galaxy-scale subhalos. Furthermore the 2-D lensing model fully exploits the numerous multiply-imaged lensing constraints available for MS2137-23 and Abell 383. The combination of gravitational lensing and stellar dynamics is the most powerful way to separate baryons and dark matter in the inner regions of clusters. However, it is important to keep a few caveats in mind. Galaxy clusters are structurally heterogeneous objects that are possibly not well-represented by simple parameterized mass models. To gain a full picture of their mass distribution and the relative contribution of their major mass components will ultimately require a variety of measurements applied simultaneously across a range of radii. Steps in this direction are already being made with the combined use of strong and weak gravitational lensing (e.g.~\citet{Limousin06,Bradac06}), which may be able to benefit further from information provided from X-ray analyses (e.g.~\citet{Schmidt06}) and kinematic studies (e.g.~\citet{Lokas03}). A recent analysis has synthesized weak-lensing, X-ray and Sunyaev-Zeldovich observations in the cluster Abell 478 -- similar cross-disciplinary work will lend further insights into the mass distribution of clusters \citep{Mahdavi07}. Of equal importance are mass models with an appropriate amount of flexibility and sophistication. For instance, incorporating models that take into account the interaction of baryons and dark matter may shed light into the halo formation process and provide more accurate representations of dark matter structure. Halo triaxiality, multiple structures along the line of sight and other geometric effects will also be important to characterize. At the moment, incorporating these complexities and securing good parameter estimates is computationally expensive even with sophisticated techniques such as the Markov-Chain Monte Carlo method. Numerical simulation results are often presented as the average profile found in the suite of calculations performed. Instead, the distribution of inner slopes would be a more useful quantity for comparison with individual cluster observations. Also, comparisons between simulations and observations would be simplified if {\it projected} density profiles of simulated halos along multiple lines of sight were to be made available. These issues should be resolvable as large samples of observed mass profiles are obtained. For the reasons above, comparing observational results with numerical simulations is nontrivial. The observational task should be regarded as one of developing mass modeling techniques of increasing sophistication that separate dark and baryonic matter, so as to provide the most stringent constraints to high resolution simulations which include baryons as they also increase in sophistication. The combination of stellar dynamics and strong lensing is the first crucial step in this process. Its diagnostic power will be further enhanced by including other major mass components (i.e.~the hot gas of the intracluster medium or the stellar contribution from galaxies) out to larger radii. A plan of the paper follows. In \S~\ref{sec:methods} we explain the methodology used to model the cluster density profile by combining strong lensing with the BCG velocity dispersion profile. In \S~\ref{sec:obsresults} we focus on translating observational measurements into strong lensing input parameters. This section includes the final strong lensing interpretation of MS2137-23 and Abell 383. In \S~\ref{sec:stronglens} we present the results of our combined lensing and dynamical analysis. In \S~\ref{sec:systematics} we discuss further systematic effects, limitations and degeneracies that our technique is susceptible to -- with an eye towards future refinements. Finally, in \S~\ref{sec:finale} we summarize and discuss our conclusions. Throughout this paper, we adopt $r$ as the radial coordinate in 3-D space and $R$ as the radial coordinate in 2-D projected space. When necessary, we assume $H_{0}$=65 \kms Mpc$^{-1}$, $\Omega_{m}$=0.3, and $\Omega_{\Lambda}$=0.7. | \label{sec:systematics} In the previous section we have presented the results of our analysis, which showed that a mass model comprising a stellar component for the BCG following a Jaffe profile together with a generalized NFW DM cluster halo is able to adequately reproduce the observations for Abell 383 (albeit {\it only} for the coarse lensing positional accuracy scenario) but is unable to simultaneously reproduce the observed multiple image configuration and BCG velocity dispersion profile for MS2137-23. In the case of Abell 383, the inner DM profile is flatter than $\beta$=1, supporting the earlier work of S04. This indicates that at least some galaxy clusters have inner DM slopes which are shallower than those seen in numerical simulations -- but only if the mass parameterization used in the current work is reflective of reality. Further work in this interesting cluster using other observational probes will further refine the mass model, and determine if the generalized NFW DM form is a good fit to the cluster profile. In this section we discuss systematic uncertainties in our method and possible refinements that could be made to reconcile the mass model with the observations for MS2137-23. We hope that many of these suggestions will become important as cluster mass models improve and thus will present fruitful avenues of research. \subsection{Systematic Errors} We focus first on systematic errors associated particularly with the troublesome stellar velocity dispersion profile for MS2137-23. Errors could conceivably arise from (i) significant non-Gaussianity in the absorption lines (which are fit by Gaussians), (ii) uncertain measurement of the instrumental resolution used to calibrate the velocity dispersion scale, and (iii) template mismatch. Non-gaussianity introduces an error that we consider too small to significantly alter the goodness of fit \citep{Gavazzi05}. The instrumental resolution of ESI (the Keck II instrument used to measure the velocity dispersion profile; \citet{Sheinis02}) is $\sim$30 \kms; this is much smaller than the measured dispersion. Even if the instrumental resolution was in error by a factor of two, the systematic shift in $\sigma$ would only be 3 \kms (using Eq~3 in Treu et al.\ 1999). This would affect all measurements and not reverse the trend with radius. Concerning template mismatch, S04 estimated a possible systematic shift of up to 15-20 \kms . This could play a role especially as the signal to noise diminishes at large radii, where the discrepancy with the model profile is greatest. To test this hypothesis, we added 20 \kms in quadrature to only those velocity dispersion data points in MS2137-23 at $R > 4$ kpc and recalculated the best-fitting $\chi^{2}$ values. $\chi^{2}$ is reduced from 31 to 28.8, a modest reduction which fails to explain the poor fit. Although selectively increasing the error bars on those data points most discrepant with the model is somewhat contrived, our result does highlight the need for high S/N velocity dispersion measures out to large radii. A high quality velocity dispersion profile has been measured locally for Abell 2199 to $\sim$20 kpc \citep{Kelsonetal02}. Interestingly, these high S/N measures display similar trends to those for MS2137-23 in the overlap regime, i.e. a slightly decreasing profile at $R\lesssim 10$kpc. The dip witnessed in MS2137-23 is thus not a unique feature, although with deeper measurements we might expect to see a rise at larger radii as a result of the shallow DM profile. A final potential limitation in the dynamical analysis is the assumption of orbital isotropy. Both S04 and \citet{Gavazzi05} explored the consequences of mild orbital anisotropy, concluding a possible offset of $\Delta \beta \sim0.15$ might result. Even including orbital anisotropy into his analysis, \citet{Gavazzi05} was unable to fit the observed velocity dispersion profile. Since we determine our lensing $\chi^{2}$ values in the source plane, we checked to make sure that no extra images were seen after remapping our best-fit lensing + velocity dispersion models back to the image plane. No unexpected images were found, although several images that were explicitly not used as constraints were found, such as the mirror image of radial arc image 2a in MS2137 and the complex of multiple images associated with 3abc, 5ab, and 6ab in Abell 383 (see Figures~\ref{fig:mulplot} and \ref{fig:mulplota383}). As discussed in \S~\ref{sec:lensinterpms2137} and \ref{sec:lensinterpa383}, some of these multiple images were not used as constraints because we could not confidently identify their position either due to galaxy subtraction residuals or blending with other possible multiple image systems. \begin{figure*} \begin{center} \mbox{ \mbox{\epsfysize=4.5cm \epsfbox{f5a.eps}} \mbox{\epsfysize=4.5cm \epsfbox{f5b.eps}} } \mbox{ \mbox{\epsfysize=4.5cm \epsfbox{f5c.eps}} \mbox{\epsfysize=4.5cm \epsfbox{f5d.eps}} } \caption{Confidence contours (68\%,95\%, and 99\%) when we allow the dark matter scale radius to be fixed at values a factor of two beyond our observationally motivated prior. Top Row -- Contours when we fix the dark matter scale radius to $r_{sc}$=50 and 400 kpc in MS2137. Although the $r_{sc}$=400 kpc scenario provides a relatively good fit to the data ($\chi^{2}\sim$26), this value for the scale radius is much larger than that observed from weak lensing data. The $r_{sc}$=50 kpc scenario is a significantly worse fit to the data, with $\chi^{2}\sim$39. Note that the DM inner slope is $\beta < 1$ in both scenarios. Bottom Row -- Contours when we fix the dark matter scale radius to $r_{sc}$=50 and 400 kpc in A383. The large discrepancy in inner slope values obtained emphasize the need for a mass probe at larger radii. The best-fitting model for either fixed scale radius is significantly worse than the best-fitting $r_{sc}$=100 kpc result ($\chi^{2}\sim$26.5 and 31.3 for $r_{sc}$=50 and 400 kpc respectively). \label{fig:diffrsc}} \end{center} \end{figure*} We finally comment on the uncertainties assigned to the multiple image systems for our lens models. We have presented two sets of results in this work; with assigned image positional accuracies of $\sigma_{I}$=0\farcs2 and 0\farcs5. We find a variety of lens models are compatible with the $\sigma_{I}$=0\farcs2 case and only when the velocity dispersion data is included into the analysis does the data fail to be reproduced by the model. Certainly if we were to further increase the positional errors, at some point a good velocity dispersion fit could conceivably be obtained, but we will refrain from doing so in the present work. Increasing the positional uncertainties is only justified if there is evidence that there are significant missing components in the mass models. Further observations that can probe the mass distribution on fine scales to larger radii and higher quality models which can account for phenomena such as adiabatic contraction in the inner regions of galaxy clusters and triaxiality represent the best way to obtain a more precise picture of the cluster mass distribution. \subsection{Improving the Mass Model} We now turn our attention to possible inadequacies in the mass model. It is important to stress that the two diagnostics (lensing and dynamics) adopted in this study probe different scales. The lensing data tightly constrains the mass profile at and outside the radial arc ($\sim$20 kpc), while the velocity dispersion constrains the mass profile inside $R \lesssim 10$ kpc. Since multiple images are numerous and their positions can be more precisely measured than velocity dispersion \footnote{The error on the astrometry with respect to the relevant scale, the Einstein Radius $\theta_{\rm E}$ is much smaller than the relative error on velocity dispersion, i.e. $\delta \theta / \theta_{\rm E} << \delta \sigma/\sigma$}, they carry more weight in the $\chi^2$ statistic than the kinematic points, producing a best overall fitting model (which is lensing dominated) that is a relatively poor fit to the kinematic data. To improve the model, one must admit that either one of the two components of the modeling is incorrect, or that the functional form of the mass profile chosen to extrapolate the lensing information at the scales relevant for dynamics is insufficient. In this section we discuss several areas where the mass model presented in this paper could be improved. \subsubsection{The Contribution of the Brightest Cluster Galaxy} We might query the assumption of a Jaffe density profile for the BCG. This seems an unlikely avenue for improvement given the Jaffe profile fits the observed BCG surface brightness profile remarkably well (see Figure 2 of S04). Moreover, \citet{Gavazzi05} utilized a Hernquist mass profile in his analysis of MS2137-23, which also matches the observations, and \citet{Gavazzi05} was likewise unable to reproduce the observed S04 velocity dispersion profile. We have additionally checked our assumptions by altering the PIEMD fit to the BCG surface brightness data so that it is matched not to the derived Jaffe profile fit to the BCG but directly to the HST surface brightness profile. With this setup, we found a $r_{cut}$ value of 23.70 kpc for MS2137 and 28.65 kpc for Abell 383 (compare this with the numbers in Table~\ref{tab:lensfixed}). Redoing our analysis for the best-fitting $r_{sc}$ scenario only, our constraints on $\beta$ for both Abell 383 and MS2137 did not change by more than 0.05, and so it is not likely that our method for constraining the BCG mass contribution is the root cause of our inability to fit the data to a BCG + gNFW cluster DM halo mass model. Conceivably the BCG may not be coincident with the center of the cluster DM halo, as has been assumed throughout this work. It is often the case that small subarcsecond off sets between BCGs and cluster DM halos are necessary to fit lensing constraints (e.g.~\citet{gps05}). There is strong evidence that the BCG is nearly coincident with the general cluster DM halo {\it in projection} from the strong lensing work presented here and by others \citep{gavazzi03,Gavazzi05}. However, an offset could be responsible for the flat to falling observed velocity dispersion profile if the BCG were actually in a less dark matter dominated portion of the cluster. Another possibility is that there are multiple massive structures along the line of sight, which would be probed by the strong lensing analysis, but not with the velocity dispersion profile of the BCG. A comprehensive redshift survey of MS2137-23 could provide further information on structures along the line of sight. \subsubsection{The Advantage of a Mass Probe at Larger Radii}\label{sec:highrad} With our presented data set, we have seen that it is difficult to constrain the DM scale radius, $r_{sc}$ because both of our mass probes are only effective within the central $\sim$100 kpc of the clusters -- within the typical DM scale radius observed and seen in CDM simulations. For this reason, the inferred DM scale radius for both Abell 383 and MS2137-23 lay at the boundary of our assumed prior range. Future work will benefit from weak lensing data, along with galaxy kinematics and X-ray data of the hot ICM which can each probe out to large clustercentric radii. Although not the focus of the current work, pinning down the correct DM scale radius will be crucial for constraining other DM mass parameters. For instance, there is a well-known degeneracy between $r_{sc}$ and the inner slope $\beta$ (e.g.~\citet{gavazzi03,Gavazzi05}). To briefly explore this, we have reran our analysis (for the coarse positioning lensing case) for both clusters with a $r_{sc}$ of 50 and 400 kpc -- factors of two beyond our chosen $r_{sc}$ prior. We show our confidence contours in Figure~\ref{fig:diffrsc}, which are noteworthy. For example in the case of MS2137-23, if we fix $r_{sc}$=50 kpc, then the best-fitting $\beta = 0.05$. However, if $r_{sc}$=400 kpc then $\beta=0.7$, more in accordance with simulations. Interestingly, the $r_{sc}$=400kpc scenario returns a better overall $\chi^{2}\sim26$ than any model with $r_{sc}$=100-200 kpc -- even though a $r_{sc}$ of 400 kpc is clearly ruled our by extant weak lensing observations. None of the other $r_{sc}$=50,400 kpc scenarios produced $\chi^{2}$ values that were comparable to those seen with $r_{sc}$=100-200 kpc. Any further knowledge of the DM scale radius would aid greatly in constraining $\beta$ and determining the overall goodness of fit of the generalized NFW DM profile to the cluster data. X-ray studies assuming hydrostatic equilibrium \citep{Allen01vir,Schmidt06} and a combined strong and weak lensing analysis \citep{Gavazzi05} have presented data on MS2137-23 to radii much larger than that probed in this study. To check that the mass model derived from data within $\sim$ 100 kpc do not lead to results at variance with published data at larger radii, we have taken the \citet{Gavazzi05} results and compared their derived mass at large radii with an extrapolation of our mass models. Examining Figure~3 from \citet{Gavazzi05} we estimate that from his weak lensing analysis a 2D projected mass enclosed between $1.6 \times 10^{14}$ and $1.1 \times 10^{15} M_{\odot}$ at $\sim 1.08$ Mpc using the cosmology adopted in this paper. Correspondingly, if we take all of the $\Delta \chi^{2}<1.0$ models using our analysis method (the coarse positional accuracy case was use) and calculate the expected 2D projected mass enclosed at 1.08 Mpc we find values between $6.9 \times 10^{14}$ and $8.4 \times 10^{14} M_{\odot}$, well within the expected range. Note that no attempt was made to extrapolate the mass {\it profiles} derived in our analysis to larger radii than the data in this paper allow, although we are acquring weak lensing data for a large sample of galaxy clusters to perform a more extensive analysis. The purpose of this consistency check is to only ensure that the masses we derive for such large radii are not too discrepant with existing analyses. The consistency check is satisfied and lends some credence to the models. \subsubsection{Dark matter baryons interactions and Triaxiality}\label{sec:ac} The central regions of DM halos can be strongly affected by the gravitational interaction with baryons during halo formation. If stars form and condense much earlier than the DM, it is expected that the baryons will adiabatically compress the DM resulting in a halo that is {\it steeper} than that of the original \citep{Blumenthal86,Gnedin04}. Alternatively, dark matter heating through dynamical friction with cluster galaxies can counteract adiabatic contraction, leading to a shallower DM profile \citep{Elzant04,Nipoti03}. The present work takes into account neither of the above scenarios, and if any baryon-DM interaction greatly changes the cluster density profile, our assumed parameterized gNFW profile may be inappropriate. Recently, \citet{Zappacosta06} have used X-ray mass measurements in the cluster Abell 2589 to conclude that processes in galaxy cluster formation serve to counteract adiabatic contraction in the cluster environment. Certainly, more observational work is needed to understand the interplay between baryons and DM in clusters, and extended velocity dispersion profiles of BCGs in conjunction with other mass tracers at larger radii could serve as the best testing ground for the interplay of dark and luminous matter. Not only is there likely significant interplay between baryons and DM in the central regions of clusters, but real galaxy clusters are certainly triaxial and, if ignored, this may lead to biased parameter estimations and discrepancies when combining mass measurement techniques that are a combination of two- and three-dimensional. Several recent studies have considered the effects of halo triaxiality on observations. Using an N-body hydrodynamical simulation of a disk galaxy and performing a 'long slit' rotation curve observation, \citet{Hayashi04} found that orientation and triaxial effects can mistake a cuspy DM profile for one that has a constant density core. At the galaxy cluster scale, \citet{Clowe04} performed mock weak lensing observations of simulated galaxy clusters and found that the NFW concentration parameter recovered was correlated with the 3D galaxy cluster orientation. In order to investigate the recent rash of galaxy clusters with observed high concentration parameters in seeming contradiction to the CDM paradigm \citep{Kneib03,Gavazzi05,Broadhurst05b}, \citet{Oguri05} used strong and weak lensing data in Abell 1689 along with a set of models that included halo triaxiality and projection effects. Again, it was seen that halo shape causes a bias in mass (and mass profile) determination, although it should be kept in mind that measurements of concentration are extremely difficult (e.g. Halkola et al.\ 2006), and the recent study of \citet{Limousin06} has seemed to clear up the concentration parameter controversy for at least Abell 1689. In terms of the current work, \citet{Gavazzi05} has pointed out that the inability of his lensing model to fit the MS2137-23 BCG velocity dispersion profile may be due to halo triaxiality or projected mass along the line of sight (which would increase the mass measured in the lensing analysis but would not be seen in the stellar velocity dispersion). \citet{Gavazzi05} showed that an idealized prolate halo with an axis ratio of $\sim$ 0.4 could explain the velocity dispersion profile in MS2137-23. Halo triaxiality could also explain the high concentration previously seen in this cluster. Again, the gap between simulations and observations may be bridged with respect to triaxiality if further steps were taken to compare the two directly. One step in this direction would be the publication of detailed density profiles for the simulations (in 3-D or along numerous projected sight-lines). The most recent DM only simulations have indicated that the standard NFW profile representation of a DM profile (and its \citet{M99} counterpart with an inner slope $\beta \sim 1.5$) can be significantly improved by slightly altering the model to a profile with a slope that becomes systematically shallower at small radii (e.g.~\citet{Navarro04}, but see \citet{Diemand05}). While we have adopted the traditional generalized NFW profile in this study, future work with parameterized models should move towards the latest fitting functions along with an implementation of adiabatic contraction as has already been attempted by \citet{Zappacosta06}. Note, however, that both \citet{Navarro04} and \citet{Diemand04} have stated that all fitted functions have their weaknesses when describing complicated N-body simulations and that when possible simulations and observations should be compared directly. | 7 | 10 | 0710.1069 |
0710 | 0710.1891_arXiv.txt | Our high time resolution observations of individual pulses from the Crab pulsar show that the main pulse and interpulse differ in temporal behavior, spectral behavior, polarization and dispersion. The main pulse properties are consistent with one current model of pulsar radio emission, namely, soliton collapse in strong plasma turbulence. The high-frequency interpulse is quite another story. Its dynamic spectrum cannot easily be explained by any current emission model; its excess dispersion must come from propagation through the star's magnetosphere. We suspect the high-frequency interpulse does not follow the ``standard model'', but rather comes from some unexpected region within the star's magnetosphere. Similar observations of other pulsars will reveal whether the radio emission mechanisms operating in the Crab pulsar are unique to that star, or can be identified in the general population. | Between 1994 and 2002 our group observed strong pulses from the Crab pulsar between 1 and 5 GHz at the VLA and Arecibo. Our Arecibo observations were designed with 2-ns time resolution, in order to test competing models of the radio emission mechanism. We initially concentrated on the MP, because it is stronger than the IP below $\sim$5 GHz in the star's mean profile, and also because giant pulses are more common at the phase of the MP at these frequencies \cite{Cordes2004}. Our results, and our data acquisition system, are described in \cite{HKWE}. \begin{figure}[ht] \vspace{-1.5in} \rotatebox{-90}{ \includegraphics[trim = 40 0 60 0, width=0.6\textwidth,clip]{EilekFig1.ps}} \vspace{-2.5in} \caption{An example of a Main Pulse, observed with 2.2-GHz bandwidth at 9 GHz, and coherently dedispersed \cite{HKWE}. The pulse seen in total intensity (upper panel, plotted with total intensity time resolution 6.4 ns) contains several short-lived microbursts. The dynamic spectrum (lower panel, plotted with resolution 102 ns and 19.5 MHz) shows that the microburst emission spans the full receiver bandwidth. In a few MPs individual, short-lived nanoshots are sparse enough in time to be separately identified (shown in \cite{HKWE, HE2007}); these examples reveal the dynamic spectrum of the nanoshots is relatively narrow. } \label{MPfig} \end{figure} \begin{figure}[ht] \rotatebox{-90}{ \includegraphics[trim = 40 0 60 0, width=0.6\textwidth,clip]{EilekFig2.ps}} \caption{An example of an Interpulse, observed with 2.2 GHz bandwidth at 9 GHz, and coherently dedispersed. The IP seen in total intensity (upper panel) typically contains 1 or 2 sub-bursts; thus it has a simpler time signature than the MP (as in the example of Figure \ref{MPfig}). The regular {\it emission bands} in the dynamic spectrum (lower panel) are not due to instrumental or interstellar effects, but are characteristic to the emission physics of the IP. Note that the secondary burst, seen in total intensity, coincides with the appearance of new band sets in the dynamic spectrum. Plotted with total intensity time resolution 51.2 ns, and dynamic spectrum resolution 104 ns and 19.5 MHz. } \label{IPfig} \end{figure} To follow up on these results, we extended our data acquisition system to higher time resolution. We went to higher frequencies (5 to 10 GHz) in order to take advantage of the 2.2-GHz bandwidths (and corresponding sub-ns time resolution) available at Arecibo. We recorded individual IPs as well as MPs, because at these frequencies strong pulses are much more common at the phase of the IP \cite{Cordes2004}. These new observations, carried out between 2003 and 2006, are reported in \cite{HE2007}. We were astonished to find that IPs are very different from MPs at these frequencies. The IP differs from the MP in polarization, time signature, spectrum and dispersion, as discussed below. All of our IP observations were taken betweeen 5 and 10 GHz. Technical limitations, as well as the scarcity of strong IPs at lower frequencies, kept us from observing the IP below 4 GHz. We are therefore describing the ``high-frequency IP'', which occurs at a slighly earlier rotation phase from the ``regular'' IP (as seen at lower radio frequencies as well as in high-energy bands; {\em e.g.}, \cite{MH1996}). This phase offset suggests that the high-frequency IP may not be related at all to the regular IP; it may come from a very different part of the magnetosphere. | We have identified two different types of coherent radio emission from the Crab pulsar, one associated with the MP, the other associated with the high-frequency IP. Two different radiation mechanisms seem to be be operating within the star's magnetosphere. In addition, the higher dispersion of the high-frequency IP suggests the signal has passed through an unusually large plasma column before leaving the pulsar. Conventional wisdom ascribes the MP to emission from the open field line region above one of the star's magnetic axes. If this is the case, the high-frequency IP is probably not simply radiation from the other magnetic pole. It is more likely to come from some unexpected part of the magnetosphere; its phase offset, relative to the regular IP \cite{MH1996}, corroborates this idea. The suggestion from \cite{Maxim} that the high-frequency IP is emitted within the closed field line region may be on the right track. | 7 | 10 | 0710.1891 |
0710 | 0710.2776_arXiv.txt | {} {The nature of \sori~(S\,Ori~J053810.1$-$023626), a faint mid-T type object found towards the direction of the young \so~cluster, is still under debate. We intend to disentangle whether it is a field brown dwarf or a 3-Myr old planetary-mass member of the cluster.} {We report on near-infrared $JHK_s$ and mid-infrared [3.6] and [4.5] IRAC/Spitzer photometry recently obtained for \sori. The new near-infrared images (taken 3.82\,yr after the discovery data) have allowed us to derive the first proper motion measurement for this object. } {The colors $(H-K_s)$, $(J-K_s)$ and $K_s$\,$-$\,[3.6] appear discrepant when compared to T4--T7 dwarfs in the field. This behavior could be ascribed to a low-gravity atmosphere or alternatively to an atmosphere with a metallicity significantly different than solar. The small proper motion of \sori~(11.0\,$\pm$\,5.9~mas\,yr$^{-1}$) indicates that this object is further away than expected if it were a single field T dwarf lying in the foreground of the \so~cluster. Our measurement is consistent with the proper motion of the cluster within 1.5\,$\sigma$ the astrometric uncertainty. } {Taking into account \sori's proper motion and the new near- and mid-infrared colors, a low-gravity atmosphere remains as the most likely explanation to account for our observations. This supports \sori's membership in \so, with an estimated mass in the interval 2--7~\mj, in agreement with our previous derivation.} | Knowledge of the initial mass function is crucial for understanding the formation processes of stars, brown dwarfs and free-floating planetary-mass objects. Whether and where there is a limit for the creation of objects by direct collapse and fragmentation of molecular clouds has become one of the major goals in the study of very young populations. Planetary-mass candidates with masses in the interval 3--13 Jovian masses (\mj) have been found in various star-forming regions (e.g., Lucas \& Roche \cite{lucas00}; Zapatero Osorio et al. \cite{osorio00}; Chauvin et al. \cite{chauvin04}; Lucas et al. \cite{lucas05}; Luhman et al. \cite{luhman05}; Jayawardhana \& Ivanov \cite{ray06}; Allers et al. \cite{allers06}; Gonz\'alez-Garc\'\i a et al. \cite{gonzalez06}; Caballero et al. \cite{caballero07}). These objects are mostly free-floating but in in a few cases appear as wide companions to young brown dwarfs or low-mass stars. \sori~(S\,Ori~J053810.1$-$023626) is the coolest free-floating, planetary-mass candidate so far reported in the literature. It was discovered by Zapatero Osorio et al$.$ (\cite{osorio02a}) and lies in the direction of the \so~cluster (352~pc and 1--8~Myr, with a best estimate at 3~Myr; Perryman et al. \cite{perryman97}; Oliveira et al. \cite{oliveira02}; Zapatero Osorio et al. \cite{osorio02b}; Sherry et al. \cite{sherry04}). The spectral type of \sori~was determined at T5.5\,$\pm$\,1.0 from molecular indices measured over near-infrared $H$- and $K$-band low-resolution spectra. Mart\'\i n \& Zapatero Osorio (\cite{martin03}) obtained an intermediate-resolution spectrum from 1.17 to 1.37~$\mu$m ($J$-band), in which the K\,{\sc i} doublet at 1.25\,$\mu$m was detected. After comparison with theoretical spectra from Allard et al. (\cite{allard01}), the authors inferred an effective temperature and surface gravity of $T_{\rm eff}$\,=\,1100\,$^{+200}_{-100}$~K and log\,$g$\,=\,3.5\,$\pm$\,0.5~cm~s$^{-2}$, in agreement with the expectations for a few megayears-old T dwarf. State-of-the-art evolutionary models (Chabrier \& Baraffe \cite{chabrier00}; Burrows et al. \cite{burrows97}; Baraffe et al. \cite{baraffe98}) yield a mass of 3\,$^{+5}_{-1}$~\mj~if \sori's very young age is finally confirmed. Burgasser et al. (\cite{burgasser04}), in contrast, have raised doubts about the low-gravity atmosphere and true cluster membership of \sori. Based on the supposed similarity of the observed spectra to field T6--T7 dwarfs, these authors argued that the S\,Ori object is ``an old, massive field brown dwarf lying in the foreground of the \so~cluster''. However, this work relied on low signal-to-noise ratio data. Better quality photometry and spectra are needed to assess the true nature of this candidate. Here we present astrometric measurements, IRAC/Spitzer data and $JHK_s$ photometry for \sori. We find that this object has colors unexpected for its spectral classification, which is measured in the range T4.5--T7 with a best estimate at T6. We ascribe this to a low gravity atmosphere, with a different metallicity being an alternative, but less likely, explanation. \begin{table*} \caption[]{Log of near-infrared observations of \sori.} \label{obslog} \centering \begin{tabular}{llcccc} \hline\hline Telescope & Instrument & Field of view & Pixel & Observing dates & Exposure time \\ & & (arcmin$^2$) & (arcsec) & & (s) \\ \hline 3.5\,m CAHA & Omega-2000 & 225 & 0.45 & 2005 Oct 22 ($JH$), 2005 Oct 25 ($K_s$) & 3600 ($J$), 6000 ($H$), 7200 ($K_s$) \\ 10\,m KeckII& NIRSPEC & 0.59 & 0.18 & 2005 Oct 26 ($JK'$) & 405 ($J$), 270 ($K'$) \\ \hline \end{tabular} \end{table*} \begin{table*} \caption[]{Near-infrared (2MASS photometric system) and IRAC/Spitzer photometry of \sori.} \label{phot} \centering \begin{tabular}{lccccccc} \hline\hline Telescope & $J$ & $H$ & $K_s$ & [3.6] & [4.5] & $H-K_s$ & $J-K_s$ \\ & (mag) & (mag) & (mag) & (mag) & (mag) & (mag) & (mag) \\ \hline 3.5 m CAHA & 19.98\,$\pm$\,0.06 & 20.07\,$\pm$\,0.07 & 19.60\,$\pm$\,0.08 & ... & ... & $+$0.48\,$\pm$\,0.11 & $+$0.38\,$\pm$\,0.10 \\ Keck\,II & 19.96\,$\pm$\,0.07 & ... & 19.58\,$\pm$\,0.07 & ... & ... & ... & $+$0.38\,$\pm$\,0.10 \\ Spitzer & ... & ... & ... & 18.62\,$\pm$\,0.30 & 17.19\,$\pm$\,0.15 & ... & ... \\ \hline \end{tabular} \end{table*} | The low proper motion of \sori~($\mu$\,=\,11.0\,$\pm$\,5.9~mas~yr$^{-1}$) makes it unlikely that it is a nearby ($\le$30~pc) T dwarf. On the one hand, we compared our measurement with the motion of 192 Hipparcos stars (Perryman et al. \cite{perryman97}) within a radius of $15^\circ$ around \so~and at a distance between 80 and 130~pc, which is the distance interval expected for \sori~if it were a field, single T6 dwarf. About 70\%~of the Hipparcos stars show larger motion than \sori, suggesting that the S\,Ori object is located farther away. On the other hand, our measurement is consistent (within 1.5~$\sigma$) with the proper motion of the O9.5V--B0.5V-type star \so~AB, which is the most massive member of the cluster of the same name. We note, however, that the relative motion of the Orion OB association is directed away from the Sun (de Zeeuw et al. \cite{zeeuw99}). This makes it very hard to detect cluster members via proper motion analysis. On the contrary, radial velocity studies can be more discriminant (Jeffries et al. \cite{jeffries06}) but the extreme faintness of \sori~prevents any accurate radial velocity measurement with current instrumentation. \subsection{Color-color diagrams} Color-color diagrams are depicted in Fig.~\ref{colcol}. To put \sori~into context, we included $JHK_s$ data of more than 100 field T-dwarfs and IRAC/Spitzer photometry of 36 field T-dwarfs compiled from the literature (Knapp et al. \cite{knapp04}; Tinney et al. \cite{tinney05}; Patten et al. \cite{patten06}; Burgasser et al. \cite{burgasser06a}; Artigau et al. \cite{artigau06}; Mugrauer et al. \cite{mugrauer06}; Leggett et al. \cite{leggett99}, \cite{leggett02}, \cite{leggett07}; Liu et al. \cite{liu07}; Luhman et al. \cite{luhman07}; Looper et al. \cite{looper07}). All near-infrared colors were conveniently transformed into the 2MASS photometric system using equations quoted in Stephens \& Leggett (\cite{stephens04}), which are appropriate for ultracool dwarfs. The location of \sori~in Fig.~\ref{colcol} is challenging since this object appears as an outlier, particularly when the $K$-band magnitude is involved. Two field dwarfs lie near it in the near-infrared color-color panels: 2MASS\,J00501994$-$3322402 (Tinney et al. \cite{tinney05}), whose photometric errors are quite large, and 2MASS\,J13243559$+$6358284 (Looper et al. \cite{looper07}). The latter object is widely discussed by Looper et al. (\cite{looper07}) in terms of binarity and low-gravity atmosphere. We applied the criterion defined by Covey et al. (\cite{covey07}, eq.~2) to distinguish objects with photometric properties deviating from the properties typical of field dwarfs and found that \sori~lies more than 2\,$\sigma$ away from the near-infrared sequence defined by the field T-type brown dwarfs. In the mid-infrared wavelengths, the photometry of \sori~deviates from the field on a 1--2\,$\sigma$ level. Covey et al.'s equation takes into account the color uncertainties of \sori~and the width of the field distribution. As compared to T4--T7 field dwarfs, \sori~presents redder $(H-K_s)$, $(J-K_s)$, and [3.6]\,$-$\,[4.5] colors and a bluer $K_s$\,$-$\,[3.6] index than expected for its spectral type ($\sim$T6). However, the $(J-H)$ index is similar to that of T5--T6 field dwarfs. The $K$-band reddish nature of \sori~is also apparent in its low-resolution $HK$ spectrum. Fig.~4 of Burgasser et al. (\cite{burgasser04}) shows these data along with the spectrum of the field T6.5 2MASS\,J10475385$+$2124234. Both spectra were obtained with similar instrumentation and are normalized to unity at 1.57~$\mu$m. While the field dwarf matches reasonably well the $H$-band region of \sori, it underestimates the flux at $K$-band, supporting the redder $(H-K_s)$ index of \sori. Burgasser et al. (\cite{burgasser04}) argued that this may be indicative of a ``lower surface gravity for \sori~relative to 2MASS\,J10475385$+$2124234''. Multiplicity cannot explain the observed photometric properties of \sori. We artificially produced near-infrared colors of L--T and T--T pairs using the absolute magnitudes provided by Liu et al. (\cite{liu06}). None of the combinations were able to reproduce our observations. The comparison of our data to theory is shown in the right panel of Fig.~\ref{colcol}. The models depicted are those of Tsuji et al. (\cite{tsuji04}), but in our analysis we also employed cloudless models by Marley et al. (\cite{marley02}) and Burrows et al. (\cite{burrows06}) obtaining similar results. The agreement between the models and the field dwarf observations is reasonably good. The great majority of the mid- and late-T dwarfs lie within the log\,$g$\,=\,4.0 and 5.5~dex tracks, as expected for ``old'' dwarfs in the solar neighborhood. This is also consistent with the recent results of the spectral fitting work by Burgasser et al. (\cite{burgasser06b}). These authors employed models by Burrows et al. (\cite{burrows06}). The near-infrared photometry of \sori~and current state-of-the-art theory of ultracool dwarfs indicate that this object may have a lower-gravity atmosphere than similarly classified T dwarfs in the solar vicinity. Because of the different pressure and density conditions at which H$_2$, CH$_4$, and CO absorptions are produced, low-gravity objects tend to be brighter at $K$ and redder in all near-infrared and [3.6]\,$-$\,[4.5] colors than comparable high-gravity objects (see discussions and Figures in Knapp et al. \cite{knapp04}; Patten et al. \cite{patten06}; Leggett et al. \cite{leggett07}; Liebert \& Burgasser \cite{liebert07}). This is what we qualitatively observe in \sori. From the right panel of Fig.~\ref{colcol} and using the solar-metallicity models by Tsuji et al. (\cite{tsuji04}), we derive log\,$g$\,$\sim$\,3.0 dex and $T_{\rm eff}$\,$\sim$\,1000--1100~K. This is consistent with previous results obtained from the spectral fitting analysis of low- and intermediate-resolution near-infrared spectra: $T_{\rm eff}$\,$\sim$\,800$^{+200}_{-100}$ K and log\,$g$\,$\sim$\,4.0\,$\pm$\,1.0 dex (Zapatero Osorio et al. \cite{osorio02a}), $T_{\rm eff}$\,$\sim$\,1100$^{+200}_{-100}$ K and log\,$g$\,$\sim$\,3.5\,$\pm$0.5 dex (Mart\'\i n \& Zapatero Osorio \cite{martin03}), respectively. These authors compared observations to theoretical data computed by Allard et al. (\cite{allard01}). Recently, Liu et al. (\cite{liu07}) have quantified the sensitivity of near-infrared spectra with $T_{\rm eff}$\,$\sim$\,700--900~K to changes in metallicity and surface gravity using a different grid of synthetic spectra by Burrows et al. (\cite{burrows06}). \sori~is brighter at $K$ relative to $J$ or $H$ by a factor of $\sim$1.4. Table~5 and Fig.~4 of Liu et al. (\cite{liu07}) suggest that log\,$g$ is thus lower than the field by about 1.0\,dex, in agreement with previous determinations. We caution that current state-of-the-art synthetic spectra do not provide detailed fits to the observed data (e.g., Burrows et al. \cite{burrows06}). Therefore, any quantitative result derived from the direct comparison of observations to models awaits further confirmation. On the contrary, for a given temperature qualitative predictions on the atmospheric relative behavior as a function of gravity and metal content can be more reliable. Metallicity might also be an issue. The \so~cluster has solar abundance ([Fe/H]\,=\,0.0\,$\pm$\,0.1~dex; Caballero \cite{caballero06}); we do not expect any metallicity effect when comparing cluster members to the field. From theoretical considerations, the effects of increasing abundance and decreasing gravity on the near-infrared spectra of cool T dwarfs are similar. From Fig.~20 of Burrows et al. (\cite{burrows06}), which shows $(J-K)$ against $T_{\rm eff}$ for various gravities and metal abundances, we infer a metallicity of [Fe/H]\,$\sim$\,$+$0.5 dex for \sori~if its reddish effect was all due to metallicity. A similar rich metal content is obtained from Liu et al. (\cite{liu07}). Nevertheless, the super-solar metallicity explanation, although possible, seems unlikely. On the one hand, the metallicity distribution of F, G, and K dwarf stars in the solar neighborhood peaks at around [Fe/H]\,=\,0.0 dex and roughly extends up to $+$0.5~dex; less than $\sim$10\%~of the stars are more metal-rich than $+$0.3 dex (Valenti et al. \cite{valenti05}; Santos et al. \cite{santos05}; Boone et al. \cite{boone06}). On the other hand, the IRAC/Spitzer data of the S\,Ori object do not support the high metallicity case. Burgasser et al. (\cite{burgasser06b}) and Liebert \& Burgasser (\cite{liebert07}) have demonstrated 2MASS\,J12373919$+$6526148 (T6.5) and 2MASS J09373487$+$2931409 (T6p) to be old, high surface gravity brown dwarfs with sub-solar abundance. These field dwarfs display [3.6]\,$-$\,[4.5] colors slightly redder (by 0.15\,mag) than expected for their assigned spectral types. In contrast to what could be inferred from the near-infrared colors, this would indicate that \sori~is a low metallicity T dwarf. Thus, a low-gravity atmosphere remains as the most likely explanation to account for the observed photometry of \sori. \begin{figure*} \centering \includegraphics[width=9cm]{4aanda.ps}~~~ \includegraphics[width=9cm]{5aanda.ps} \caption{Color-magnitude diagram of \so~low mass members. \sori~is labeled. The 3-Myr isochrone by Chabrier \& Baraffe (\cite{chabrier00}) is overplotted onto the data as a solid line in the left panel (see text for $T_{\rm eff}$, log\,$L/L_\odot$ conversion into observables). The right panel shows the sequence of M5.5--T8 field dwarfs at the distance of the cluster (solid line, spectral types are indicated). The dotted line stands for the field sequence shifted by $-$1.8 mag in the $J$-band to match the photometric trend delineated by \so~members. Objects with $J$\,$-$\,[3.6] colors significantly redder than the cluster sequence show infrared flux excesses likely due to circum(sub)stellar disks (Caballero et al. \cite{caballero07}).} \label{colmag} \end{figure*} No obvious infrared flux excesses are detectable in the IRAC/Spitzer [3.6] and [4.5] bands, suggesting that there is no envelope or disk around \sori~emitting intensively at these wavelengths. A possitive detection would have provided strong evidence for its youth. However, we remark that disks around young, low-mass brown dwarfs (close to the deuterium burning mass limit) are seen at wavelengths longer than 5\,$\mu$m (Luhman et al. \cite{luhman05}; Caballero et al. \cite{caballero07}; Zapatero Osorio et al. \cite{osorio07}), while the observed fluxes in the near-infrared up to 5\,$\mu$m are photospheric in origin. The public [5.8]- and [8.0]-band IRAC/Spitzer images are not conclusive for \sori. \subsection{Color-magnitude diagram} The photometric sequence of \so~substellar members, including \sori, is shown in the $J$ vs $J$\,$-$\,[3.6] color-magnitude diagram of Fig.~\ref{colmag} (Caballero et al. \cite{caballero07}; Zapatero Osorio et al. \cite{osorio07}). This sequence follows a relatively smooth progression with increasing color down to $J$\,$\sim$20 and $J$\,$-$\,[3.6]\,=\,2.8~mag. The location of \sori~suggests that the $J$\,$-$\,[3.6] index suddenly turns toward bluer values at a nearly unchanged $J$ magnitude. A similar turnover (ocurring at spectral types L7--L8, 1400--1300~K) is also observed in field ultracool dwarfs. The field sequence of objects with spectral types M5.5--T8 moved to the distance of the \so~cluster is displayed in the right panel of Fig.~\ref{colmag} (absolute magnitudes and colors are adopted from Patten et al. \cite{patten06}, and references therein). Because of their very young age, \so~low-mass stars and substellar objects are in the phase of gravitational contraction (e.g., Chabrier \& Baraffe \cite{chabrier00}). Cluster members thus show larger size and luminosity than their older counterparts of related colors in the field. As seen in Fig.~\ref{colmag}, the average cluster photometric sequence appears brighter than the field by about 1.8~mag in the $J$-band. The dotted line in Fig.~\ref{colmag} (right panel) represents the field sequence normalized to the \so~locus of late-M and early-L cluster members. \sori~nicely sits on the location expected for \so~T-type members. We have also compared our data to the {\sc cond} and {\sc dusty} solar metallicity evolutionary models by Chabrier \& Baraffe (\cite{chabrier00}). For the age range 1--8\,Myr, substellar objects with $T_{\rm eff}$ between 700 and 1300~K, corresponding to the mass interval $\sim$1--6\,\mj, show log\,$g$\,=\,3.0--4.0 dex, which coincides within the large error bar with the surface gravity estimation for \sori. The 3-Myr isochrone (Chabrier \& Baraffe \cite{chabrier00}) is displayed along the photometric sequence of \so~in the left panel of Fig.~\ref{colmag}. Theoretical surface temperatures and luminosities were converted into observed magnitudes and colors using the color-temperature-spectral type and spectral type-bolometric correction relationships given in the literature (Dahn et al. \cite{dahn02}; Vrba et al. \cite{vrba04}; Knapp et al. \cite{knapp04}; Patten et al. \cite{patten06}). The model convincingly reproduces the cluster low mass sequence except for the fact that \sori~appears overluminous by about 1~mag, which might suggest binarity. This was also discussed in Zapatero Osorio et al. (\cite{osorio02a}). However, there are issues that prevent us from concluding whether this object is double or whether models make wrong predictions for the smallest masses and young ages: {\sl (i)} as seen from the field sequence (Fig.~\ref{colmag}, right panel), there is a $J$-band brightening across the color turnover that theory fails to reproduce (Vrba et al. \cite{vrba04}; Knapp et al. \cite{knapp04}). {\sl (ii)} The blue color turnover takes place at a roughly constant temperature in the field ($\sim$1300--1400~K), and all color-temperature-bolometric correction transformations show a sharp change at this point; on the contrary, the evolutionary models available to us do not have a complete temperature sampling (this may explain the abrupt color reversal at the bottom of the isochrone in Fig.~\ref{colmag}). {\sl (iii)} The relations used to transform theoretical predictions into observables are obtained for high-gravity field objects. It is now known that gravity impacts significantly the near- and mid-infrared colors of T dwarfs (Leggett et al. \cite{leggett07}; Burrows et al. \cite{burrows06}), whereas the colors of the warmer M and L types are not so sensitive to the gravity parameter. Transformations are thus expected to be gravity-dependent for the coolest temperatures. It becomes necessary to find a physical explanation for the $J$ brightening feature and to discover more \so~T-type, planetary-mass members for a proper comparison with evolutionary tracks. | 7 | 10 | 0710.2776 |
0710 | 0710.2295_arXiv.txt | I discuss the development and resolution of the solar neutrino problem, as well as opportunities now open to us to extend our knowledge of main-sequence stellar evolution and neutrino astrophysics. | } This paper is based on a talk given at the Caltech conference \cite{caltech} ``Nuclear Astrophysics 1957-2007: Beyond the First 50 Years," July 23-27, 2007, which focused on the state of nuclear astrophysics fifty years after the seminal paper of Burbidge, Burbidge, Fowler, and Hoyle \cite{bbfh}. The quest to measure solar neutrinos, and later to resolve the solar neutrino problem, began in the early days of nuclear astrophysics, with the first efforts to understand proton burning in main sequence stars. I would like to review that history, our current understanding of solar neutrinos, and open questions in neutrino physics, and discuss some opportunities for further solar neutrino measurements. Solar neutrino physics brings together stellar modeling, nuclear reactions, and observation. A key early development was the 1959 Holmgren and Johnston \cite{holmgren} measurement of the S-factor for the pp-chain reaction $^3$He($\alpha,\gamma)^7$Be, which proved to be surprisingly large. This implied that the sun could produce some of its energy through the more temperature dependent ppII and ppIII cycles of the pp chain, elevating the neutrino fluxes expected from $^7$Be electron capture and $^8$B $\beta$ decay. (See Fig. 1.) These neutrinos contribute to ground- and excited-state transitions in $^{37}$Cl($\nu$,e)$^{37}$Ar, a reaction for detecting neutrinos that had been proposed by Pontecorvo \cite{pontecorvo} in 1946 and considered by Alvarez \cite{alvarez} in 1949 (who studied backgrounds that might inhibit, for example, a reactor neutrino experiment). Alvarez's envisioned experiment -- proposed as a test of the distinguishability of the neutrino and antineutrino, prior to the discovery of parity violation -- was later conducted by Davis at Savannah River \cite{jandr,davis55}. \begin{figure} \begin{center} \includegraphics[width=12cm]{fig1.pdf} \end{center} \caption{The three cycles of the pp chain and associated neutrinos.} \label{fig1} \end{figure} The Caltech effort in nuclear astrophysics brought three young researchers, Bahcall, Iben, and Sears, together in the summer of 1962. Stimulated in part by the Holmgren and Johnston measurement, they began construction of a solar model to predict the solar core temperature, the most important parameter governing the competition between the ppI, ppII, and ppIII cycles, and to provide the first quantitative estimate of the resulting neutrino fluxes. The Bahcall, Fowler, Iben, and Sears model, published in 1963, predicted a counting rate for Davis's proposed 100,000-gallon chlorine solar neutrino detector of about one event per day. A key development occurred in 1963 when, during a seminar by Bahcall at Copenhagen, Mottelson inquired about the importance of neutrino excitation of excited states in $^{37}$Ar \cite{jandr}. In 1964 Bahcall and Barnes \cite{barnes} pointed out that a calibration of the excited state contribution to the $^{37}$Cl cross section could be made by measuring the delayed protons from the analog $\beta$ decay of $^{37}$Ca, the isospin mirror of $^{37}$Cl. Effectively the lifetime of $^{37}$Ca would be a test of the elevated cross section predicted on the basis of the excited-state contribution to $^8$B neutrino absorption. Subsequent measurements by Hardy and Verrall \cite{hardy} and Reeder, Poskanzer, and Esterlund \cite{poskanzer} established the importance of the excited-state contribution. [These experiments were later repeated in a manner that was kinematically complete: see Adelberger et al. \cite{adelberger} for a discussion.] Bahcall \cite{bahcall64} and Davis \cite{davis64} published companion letters in March 1964 arguing the adequacy and feasibility of a 100,000-gallon Cl experiment to measure solar neutrinos. Excavation of the detector cavity in the Homestake Mine began in summer 1965. The tank was installed and filled by the next year. The first results were announced by Davis, Harmer, and Hoffman \cite{davis68} in 1968, an upper bound of 3 SNU (1 SNU = 10$^{-36}$ capures/Cl atom/sec), below the standard solar model (SSM) prediction of Bahcall, Bahcall, and Shaviv of 7.5 $\pm$ 3 SNU \cite{bbs}. This result and associated theoretical work on suggested solutions led to a series of experiments -- Gallex/GNO/SAGE \cite{gallex,gno,sage}, Kamiokande \cite{kamiokande}, Super-Kamiokande \cite{sk} , and the Sudbury Neutrino Observatory \cite{sno}. Efforts by Borexino \cite{borexino} and KamLAND \cite{kamland} to measure the $^7$Be neutrinos are currently underway or in preparation. These experiments are important as tests of the SSM and of the new neutrino physics that proved to be the source of the solar neutrino problem. | Neutrino astrophysics and the theories of the origin of the elements, the main theme of this conference, share a common history. Laboratory astrophysics has made solar neutrino physics into a quantitative field, and allowed experimenters to anticipate the kinds of major discoveries that justified experiments like SNO and Super-Kamiokande. The results -- discovery of neutrino mass and flavor mixing characterized by large angles -- are of great importance, providing our first constraints on physics beyond the SM of particle physics. But as summarized here, the list of remaining laboratory neutrino physics questions is long. The answers to the open questions will be important in helping us characterize extreme astrophysical and cosmological neutrino environments. The needed 20-year program of laboratory and astrophysical neutrino studies is not unlike the laboratory/astrophysics interface that Willie Fowler cultivated to help us understand the origin of the elements. Despite the current focus on particle physics properties of neutrinos, solar neutrino spectroscopy remains an important probe of the SSM and stellar evolution. The arguments for measuring the CNO flux, using our sun as a calibrated laboratory, seem particularly strong. Such a program would effectively test our understanding of the hydrogen burning mechanism for massive main-sequence stars. It would also address the primary discrepancy in the SSM, the tension between helioseismology and neutrino flux predictions that follows from new analyses of surface metallicity. | 7 | 10 | 0710.2295 |
0710 | 0710.2892_arXiv.txt | The recently recognized class of ``transitional disk" systems consists of young stars with optically-thick outer disks but inner disks which are mostly devoid of small dust. Here we introduce a further class of ``pre-transitional disks" with significant near-infrared excesses which indicate the presence of an optically thick inner disk separated from an optically thick outer disk; thus, the spectral energy distributions of pre-transitional disks suggest the incipient development of disk gaps rather than inner holes. In UX Tau A, our analysis of the {\it Spitzer} IRS spectrum finds that the near-infrared excess is produced by an inner optically thick disk and a gap of $\sim$56 AU is present. The {\it Spitzer} IRS spectrum of LkCa 15 is suggestive of a gap of $\sim$46 AU, confirming previous millimeter imaging. In addition, UX Tau A contains crystalline silicates in its disk at radii $\gtrsim$ 56 AU which poses a challenge to our understanding of the production of this crystalline material. In contrast, LkCa 15's silicates are amorphous and pristine. UX Tau A and LkCa 15 increase our knowledge of the diversity of dust clearing in low-mass star formation. | Previous studies have revealed stars with inner disks that are mostly devoid of small dust, and these ``transitional disks" have been proposed as the bridge between Class II objects, young stars surrounded by full disks accreting material onto the central star, and Class III objects, stars where the protoplanetary disk is mostly dissipated and accretion has stopped (e.g. Strom et al. 1989, Skrutskie et al. 1990; Stassun et al. 2001). New spectra from the {\it Spitzer Space Telescope} which greatly improve our resolution in the infrared have been used to define the class of ``transitional disks" as those with spectral energy distributions (SEDs) characterized by a significant deficit of flux in the near-infrared relative to optically thick full disks, and a substantial infrared excess in the mid- and far-infrared. Extensive modeling studies of several transitional disks around T Tauri stars \citep{dalessio05, uchida04, calvet05, espaillat07} and F-G stars \citep{brown07} have been presented. In particular, the SEDs of the transitional disks of the T Tauri stars (TTS) CoKu Tau$/$4 \citep{dalessio05}, TW Hya \citep{calvet02, uchida04}, GM Aur, DM Tau \citep{calvet05}, and CS Cha \citep{espaillat07} have been explained by modeling the transitional disks with truncated optically thick disks with most of the mid-infrared emission originating in the inner edge or ``wall" of the truncated disk. In all these cases, except in CoKu Tau$/$4, material is accreting onto the star, so gas remains inside the truncated disk, but with a small or negligible amount of small dust, making these regions optically thin. Here we present models of UX Tau A and LkCa 15, low-mass pre-main sequence stars in the young, $\sim$1 Myr old Taurus star-forming region which have been previously reported as transitional disks \citep{furlan06, bergin04}. We present evidence for gaps in optically thick disks, as opposed to ``inner holes", that is, large reductions of small dust from the star out to an outer optically thick wall. | Here we introduce the ``pre-transitional disk" class where we see the incipient development of disk gaps in optically thick protoplanetary disks as evidenced by significant near-infrared excesses when compared to the Taurus median SED and previously studied transitional disks \citep{dalessio05, calvet02, uchida04, calvet05, espaillat07}. The pre-transitional disk of UX Tau A has a $\sim$56 AU gap as opposed to an inner hole. It is also possible to fit LkCa 15's SED with a $\sim$46 AU gap that contains some optically thin dust; a model that has a hole rather than a gap also fits its SED and future near-infrared interferometry may be able to discriminate between these models. However, the truncation of LkCa 15's outer disk at $\sim$46 AU is consistent with resolved millimeter interferometric observations \citep{pietu06} which makes it one of three inner disk holes imaged in the millimeter (TW Hya: Hughes et al. 2007; GM Aur: Wilner et al. in preparation). In addition to our sample, the disks around F-G stars studied by \citet{brown07} also belong to the pre-transitional disk category. The large gaps that are being detected in pre-transitional disks are most likely due to observational bias since larger gaps will create larger mid-infrared deficits in the SED. Smaller gaps will most likely have less apparent dips in their SEDS and be more difficult to identify, however, if their gaps contain some optically thin material the silicate emission in these objects should be much stronger than can be explained by a full disk model. The existence of an inner optically thick disk may be an indicator of the first stages of disk clearing that will eventually lead to the the inner holes that have been seen in previously reported transitional disks; this has important implications on disk evolution theories since only planet-formation can account for this structure. Hydrodynamical simulations have shown that a newly formed planet could accrete and sweep out the material around it through tidal disturbances and this is sufficient in producing the hole size in CoKu Tau$/$4 \citep{quillen04}, even maintaining substantial accretion rates \citep{varniere06}. Moreover, \citet{najita07} have found that the intrinsic properties of transitional disks may favor planet formation. Another proposed formation mechanism for the holes in transitional disks is photoevaporation, in which a photoevaporative wind halts mass accretion towards the inner disk and material in this inner disk is rapidly evacuated creating an inner hole \citep{clarke01}; the hole then increases in size as the edge continues photoevaporating \citep{alexanderarmitage}. Neither this model nor the inside-out evacuation induced by the MRI \citep{chiang07} would explain how an optically thick disk accreting at a sizable accretion rate (see Table 1) would remain inside the hole. Rapid dust growth and settling has also been proposed to explain the holes in disks \citep{lin04}. Again, this does not account for the presence of optically thick inner disk material given that theory suggests grain growth should be fastest in the inner disk, not at some intermediate radius \citep{weiden97, chiang99}. Our sample also has interesting dust compositions (Watson et al. 2007, Sargent et al. in preparation). LkCa 15 has an amorphous silicate feature indicating little if any processing leading to the crystallization seen in other young stars. Amorphous silicates are also seen in CoKu Tau$/4$, DM Tau, and GM Aur. In contrast, UX Tau A is different from all the other transitional disks because it has crystalline silicate emission features in addition to amorphous silicate emission features (Fig. 1 inset). The wall at $\sim$56 AU is the main contributor to the crystalline silicate emission since it dominates the flux in the mid- and far-infrared. This raises the question of whether crystalline silicates are created close to the star or if they can be created in situ at $\sim$56 AU. If the former, it challenges current radial-mixing theories, none of which can get significant amounts of crystalline silicates out to this distance \citep{gail01, kellergail04}. One possibility for {\it in situ} processing may be collisions of larger bodies, which might produce small grains heated sufficiently to create crystals (S. Kenyon, personal communication). Pre-transitional disks offer further insight into the diversity of the ``transitional disk" class and future studies of these disks will greatly advance our understanding of disk evolution and planet formation. \vskip -0.1in | 7 | 10 | 0710.2892 |
0710 | 0710.0545_arXiv.txt | We present a model to estimate the synchrotron radio emission generated in microquasar (MQ) jets due to secondary pairs created via decay of charged pions produced in proton-proton collisions between stellar wind ions and jet relativistic protons. Signatures of electrons/positrons are obtained from consistent particle energy distributions that take into account energy losses due to synchrotron and inverse Compton (IC) processes, as well as adiabatic expansion. The space parameter for the model is explored and the corresponding spectral energy distributions (SEDs) are presented. We conclude that secondary leptonic emission represents a significant though hardly dominant contribution to the total radio emission in MQs, with observational consequences that can be used to test some still unknown processes occurring in these objects as well as the nature of the matter outflowing in their jets. | \label{intro} X-ray binary systems (XRBs) are composed by either a stellar mass black hole or a neutron star, and a normal (non degenerated) star which supplies matter to the compact object through the formation of an accretion disk. Some ~260 XRBs are known up to now \cite{liu06} probably corresponding to an underlying population of some tens of thousands of compact objects in our Galaxy. A few of these sources present also non-thermal radio emission, hence evidencing the existence of mechanism(s) capable of injecting and/or accelerating large numbers of relativistic particles. Some radio emitting X-ray binary systems (REXBs) have been observed showing ejection of material at relativistic velocities and to display jets like those seen in quasars and active galactic nuclei but at $\sim 10^{-6}$ times shorter scales. This analogy is the reason for calling them microquasars (MQs) \cite{mirabel99} and making them some of the most interesting objects for astrophysics. Furthermore, attention on these objects has grown since the proposal of Paredes et al. (2000) \cite{paredes00} of MQs as counterparts of some of the unidentified gamma-ray sources of the EGRET catalog \cite{hartman99} and hence pointing them as plausible high energy emitters. A strong confirmation of this association has come from the detections of the MQs LS 5039 and LS I +61 303 at Tev energies using respectively the ground-based Cherenkov telescopes HESS \cite{aharonian05} and MAGIC \cite{albert06}, giving support and empowering at the same time a number of previous detailed studies centered on the mechanisms operating in these sources in order to explain the gamma ray domain (see, e.g., \cite{bosch05} and \cite{romero05}). Moreover, a jet origin of the emission from MQs has been suggested from the observation of syncrothron emission of relativistic electrons/positrons extending from the radio all the way into the X-ray regime. In this sense jet-like models have focused on different approaches regarding the particle origin that could generate the required emission properties in a consistent way. Some of them consider leptons directly injected at the base of the jet and, in extending outwards, Compton-interaction with external/self-created photon fields produces high energy radiation. Other models deal with an hadronic origin of the high energy emission, through proton-proton interactions and pion decay producing gamma rays and leaving the resulting co-generated leptons as low energy emitters. The present work refers to the later procedure, focusing on the modelisation of the secondary leptonic synchrotron emission in order to constrain the characterization of MQ jets. An outline of the model is given in the next, followed by the results showing the SEDs and lightcurves under different parameter assumptions and the conclusions that can be extracted from them. | SEDs are obtained for different magnetic field values, electron/positron spectral indices and spatially distributed disks. We have estimated also the expected emission along the jet at 1 and 5 GHz. Leptons are injected in the context of hadronic secondaries generation within a detailed model that takes into account in a consistent way particle injection mechanisms and cooling due to radiation processes and adiabatic expansion. The luminosities obtained are slightly lower than in the models based on primary leptons injection, and must be considered complementary to them. However, we note that within our model there is no requirements of acceleration processes along the jet to obtain the final emission results. Such acceleration processes are still not well understood, although. They could come from diffusive shock acceleration along the jet when fresh ejecta interact with previous blobs of plasma already outflowing at lower velocities. Other scenarios assume a continuous energy transfer mechanism from the magnetic field to the matter content of the jet in such a way that the resulting parsec-scale radio emission can be explained. The fact of studying alternative models were particles are directly injected until a certain height along the jet can constrain the amount of acceleration required and contribute to the understanding of the physical mechanisms that can lead to such processes. Signatures at different distances along the jet and specific spectral features detectable for reasonable parameter values treated in our numerical simulations have the potential to be an important clue for determining the matter content of jets. In particluar, highly resolved observations at 1 and 5 GHz could determine if leptons are present at heights $10^{12-13}$ cm at the edge of the binary system typical region where wind matter from the companion is still significant. If electrons/positrons still show high energies due to a recent injection from hadronic interactions at these parts of the jet, it could be a signature of secondary generation without the necessity of invoking additional acceleration processes. | 7 | 10 | 0710.0545 |
0710 | 0710.5010_arXiv.txt | Recent observations of XTE J1739-285 suggest that it contains a neutron star rotating at 1122 Hz\cite{Kaaret2007}. Such rotational frequency would be the first for which the effects of rotation are significant. We study the consequences of very fast rotating neutron stars for the potentially observable quantities as stellar mass and pulsar period. | \label{sect:introd} Neutron stars with their very strong gravity can be very fast rotators. Theoretical studies show that they could rotate at sub-millisecond periods, i.e., at frequency $f=1/{\rm period}>$1000 Hz\cite{CST1994,Salgado1994}. The first millisecond pulsar B1937+21, rotating at $f=641$ Hz\cite{Backer1982}, remained the most rapid one during 24 years after its detection. In January 2006, discovery of a more rapid pulsar J1748-2446ad rotating at $f=716$ Hz was announced \citep{Hessels2006}. However, such sub-kHz frequencies are still too low to significantly affect the structure of neutron stars with $M>1M_\odot$ \cite{STW1983}. Actually, they belong to a {\it slow rotation} regime, because their $f$ is significantly smaller than the mass shedding (Keplerian) frequency $f_{\rm K}$. Effects of rotation on neutron star structure are then $\propto (f/f_{\rm K})^2\ll 1$. Rapid rotation regime for $M>1M_\odot$ requires submillisecond pulsars with supra-kHz frequencies $f>1000$ Hz. Very recently Kaaret et al.\cite{Kaaret2007} reported a discovery of oscillation frequency $f=1122$ Hz in an X-ray burst from the X-ray transient, XTE J1739-285. According to Kaaret et al.\cite{Kaaret2007} "this oscillation frequency suggests that XTE J1739-285 contains the fastest rotating neutron star yet found". If confirmed, this would be the first detection of a sub-millisecond pulsar (discovery of a 0.5 ms pulsar in SN1987A remnant announced in January 1989 was withdrawn one year later). Rotation at $f>1000$ Hz is sensitive to the stellar mass and to the equation of state (EOS). Hydrostatic, stationary configurations of neutron stars rotating at given rotation frequency $f$ form a one-parameter family, labeled by the central density. This family - a curve in the mass - equatorial radius plane - is limited by two instabilities. On the high central density side, it is instability with respect to axi-symmetric perturbations, making the star collapse into a Kerr black hole. The low central density boundary results from the mass shedding from the equator. In the present paper we show how rotation at $f>1000$ Hz is sensitive to the EOS, and what constraints on the EOS of neutron stars result from future observations of stably rotating sub-millisecond pulsars. | The $M(R_{\rm eq})$ curve for $f\gtrsim 1400$ Hz is flat. Therefore, for given EOS the mass of NS is quite well defined. Conversely, measured mass of a NS rotating at $f\gtrsim 1400$ Hz will allow us to unveil the actual EOS. The "Newtonian" formula for the Keplerian frequency works surprisingly well for precise 2-D simulations and sets a firm upper limit on $R_{\rm eq}$ for a given $f$. Finally, observation of $f\gtrsim 1200$ Hz sets stringent limits on the initial mass of the nonrotating star which was spun up to this frequency by accretion. \label{sect:discuss} | 7 | 10 | 0710.5010 |
0710 | 0710.0621_arXiv.txt | X-ray spectra from stellar coronae are reprocessed by the underlying photosphere through scattering and photoionization events. While reprocessed X-ray spectra reaching a distant observer are at a flux level of only a few percent of that of the corona itself, characteristic lines formed by inner shell photoionization of some abundant elements can be significantly stronger. The emergent photospheric spectra are sensitive to the distance and location of the fluorescing radiation and can provide diagnostics of coronal geometry and abundance. Here we present Monte Carlo simulations of the photospheric K$\alpha_1,\alpha_2$ doublet arising from quasi-neutral Fe irradiated by a coronal X-ray source. Fluorescent line strengths have been computed as a function of the height of the radiation source, the temperature of the ionising X-ray spectrum, and the viewing angle. We also illustrate how the fluorescence efficiencies scale with the photospheric metallicity and the Fe abundance. Based on the results we make three comments: (1) fluorescent Fe lines seen from pre-main sequence stars mostly suggest flared disk geometries and/or super-solar disk Fe abundances; (2) the extreme $\approx 1400$~m\AA\ line observed from a flare on V~1486~Ori can be explained entirely by X-ray fluorescence if the flare itself were partially eclipsed by the limb of the star; and (3) the fluorescent Fe line detected by {\it Swift} during a large flare on II~Peg is consistent with X-ray excitation and does not require a collisional ionisation contribution. There is no convincing evidence supporting the energetically challenging explanation of electron impact excitation for observed stellar Fe~K$\alpha$ lines. | \label{s:intro} It is well-established from surveys of the sky at EUV and X-ray wavelengths that all stars with spectral types later than mid-F, except for giants later than mid-K, possess hot outer atmospheres akin to that of the Sun \citep[e.g.][]{Vaiana.etal:81, Schmitt:97}. While much observational and theoretical effort has been devoted to understanding solar coronal spectra and, in more recent years, toward understanding stellar coronal emission and spectra, comparatively little attention has been devoted to the reprocessing and line fluorescence resulting from this coronal emission by the underlying solar and stellar photospheres. In contrast, considerable effort has been spent on the study of X-ray reprocessing by ``cold'' gas in much more complex systems with more prominent fluorescent features but more uncertain geometries and physical conditions, such as the accretion disks around black holes and non-degenerate objects in X-ray binaries \citep[e.g.][]{Felsteiner.Opher:76,Hatchett.Weaver:77,Fabian.etal:89, George.Fabian:91,Laor:91, Matt.etal:97, Ballantyne.etal:02, Beckwith.Done:04, Cadez.Calvani:05, Dovciak.etal:04, Laming.Titarchuk:04, Brenneman.Reynolds:06}. The processes involved in photospheric fluorescence by coronal irradiation are the same as those discussed in these works; the main difference here is in the specific geometry of the X-ray source above a quasi-neutral sphere and of the extended, shell-like nature of the coronal source above the photosphere during quiescent conditions. X-rays emitted from a hot ($T\ga 10^6$~K) corona incident on the underlying photosphere can undergo either Compton scattering or photoabsorption events through the ionization of atoms or weakly ionized species. Through scattering events, photons can be reflected back in a direction towards the stellar surface where they have a finite chance of escape. Compton scattering redistributes the spectrum to lower energies by $\sim E^2/m_ec^2$ per collision, where $E$ is the photon energy and $m_e$ the electron rest mass. The spectrum reflected from a stellar surface by scattering events is then shifted and broadened towards lower energies. Photoionization events involving X-ray photons directed toward the photosphere are predominantly inner shell interactions with astrophysically abundant elements, the outer and valance cross-sections being very small at these energies. Observable fluorescent lines can then arise as a result of the finite escape probabilities of photons emitted in outward directions by hole transitions in these atoms photoionized in their inner shells. These processes have been described in the solar context by, e.g., \citet{Tomblin:72} and \citet{Bai:79} (B79), and more recently for arbitrarily photoionized slabs by \citet{Kallman.etal:04}. The strongest of the fluorescent lines for a plasma of approximately cosmic composition is the $2s$-$1p$ 6.4~keV Fe K$\alpha$ doublet occurring following ejection of a $1s$ electron. It has been observed in solar spectra on numerous occasions \citep[e.g.][]{Neupert.etal:67, Doschek.etal:71, Feldman.etal:80, Tanaka.etal:84, Parmar.etal:84, Zarro.etal:92}. The mechanism of fluorescence by the thermal X-ray coronal continuum was suggested by \citet{Neupert.etal:67}, and was firmly established on more theoretical grounds by \citet{Basko:78,Basko:79} and B79. \citet{Parmar.etal:84} provided convincing observational confirmation based on flare spectra obtained by the {\it Solar Maximum Mission}, though it has also been noted that contributions from non-thermal electron impact might also be present during hard X-ray bursts \citep*[e.g.][]{Emslie.etal:86, Zarro.etal:92}. B79 pointed out that, for a given source spectrum, the observed flux of Fe~K$\alpha$ photons from the photosphere depends on essentially three parameters: the photospheric iron abundance; the height of the emitting source; and the heliocentric angle between the emitting source and observer. \citet{Phillips.etal:94} used Fe K$\beta$ observations to probe the difference between the photospheric and coronal iron abundance for flares observed by the {\it Yohkoh} satellite. More importantly for the stellar case, the spatial aspects of photospheric fluorescent line formation suggest its application to understanding the spatial distribution of coronal structures and flares on stars of different spectral type and activity level to the Sun \citep{Drake.etal:99}. Indeed, Fe~K fluorescence has recently been detected during flares on the active binary II~Peg \citep{Osten.etal:07} and on the single giant HR~9024 \citep{Testa.etal:07}. The Fe~K line has also been seen in a growing sample of pre-main sequence (PMS) stars \citep{Imanishi.etal:01,Favata.etal:05,Tsujimoto.etal:05, Giardino.etal:07}, in which the line is thought to originate predominantly from the irradiated protoplanetary disk rather than the photosphere. Since the work of \cite{Bai:79}, there have been no concerted efforts to extend models of photospheric fluorescence for coronal excitation sources with other characteristics. Fluorescent lines other than Fe~K$\alpha$ have also not yet, to our knowledge, been studied by other workers in any detail in this context. A reasonably strong feature observed in solar spectra near 17.62~\AA\ had been identified with Fe L$\alpha$ photospheric fluorescence \citep{McKenzie.etal:80, Phillips.etal:82}, but calculations of the expected line strength was shown by \citet{Drake.etal:99} to be much too weak to explain the feature, and these authors instead identified the line with a transition in Fe~XVIII arising from configurational mixing and both seen in {\it Electron Beam Ion Trap} spectra and predicted by theory \citep{Cornille.etal:92}. However, given the potential diagnostic value of photospheric fluorescence, other lines, such as O~K$\alpha$, are possibly observable with very high quality observations and warrant further study. The capabilities of current X-ray missions such as {\it Chandra}, {\it XMM-Newton}, {\it Swift} and {\it Suzaku} to detect fluorescent lines further motivates a re-examination of the photospheric fluorescence problem in the context of stellar coronae, photospheres and protoplanetary disks. We restrict the study in hand to Fe~K fluorescence from stellar photospheres and defer detailed discussions of protoplanetary disks and fluorescence from other elements to future work. | The main scientific motivation for this work is to provide the foundation to use Fe fluorescence as a quantitative diagnostic of coronal and flare geometry. There now exists a handful of detections of fluorescent emission from stars. Sensitivity is currently limited to a large extent by the low spectral resolution of available instruments and progress is expected to accelerate dramatically with the future availability of X-ray calorimeters. Since modern X-ray spectral analyses based on low-resolution CCD pulse-height spectra tend to express line strengths in terms of the line equivalent width, we have computed this quantity for the case of $\theta=0$ and the ranges of heights and X-ray temperatures investigated in \S\ref{s:newcalcs}. The equivalent width in this context refers to the fluorescent, processed line seen on top of the continuum of the ionising coronal spectrum. While the $\theta=0$ case gives the most optimistic line strength, we note that $f(\theta)$ is quite slowly varying for angles $\theta \la 45^\circ$ for the flare heights for which significant Fe\,K${\alpha}$ might be observed. The equivalent widths are illustrated in Figure~\ref{f:ew} and listed in Table~\ref{t:fekaEW1}. \subsection{Fluorescence from Pre-Main Sequence Stars} The observability of the cold Fe~K$\alpha$ line is of course strongly dependent on the quality of the X-ray spectrum obtained. The most extensive study of PMS Fe fluorescence to date is that based on Chandra observations of the Orion Nebula Cluster by \citet{Tsujimoto.etal:05}. This study detected significant 6.4~keV excesses attributable to Fe fluorescence for 7 out of 127 sources found to have significant counts in the 6-9~keV band. Equivalent widths were in the range 110-270~eV at plasma temperatures of $\sim 3$-10~keV. There is clearly a strong selection effect here and these fluorescent line strengths likely represent the upper end of the distribution. Our calculations for a flare at scale height $h=0 R_\star$ are also appropriate for a flare occurring above an infinite plane, such as might approximate a disk-encircled PMS star. As in the photospheric case, the fluorescence problem can be treated orthogonally from the ionisation structure of the disk, which is not greatly altered from its very largely neutral overall state by X-rays from a typical flare. Any small degree of X-ray photoionisation will also not affect Fe~K$\alpha$ line strengths because fluorescence yields are essentially invariant for lower Fe ions. Our calculations indicate that attaining equivalent widths much in excess of 100~eV is not straightforward for such a simple geometry for the plasma temperatures observed in the fluorescing X-ray spectra. This finding is in agreement with earlier calculations by \citet{Matt.etal:91} and \citet{George.Fabian:91}, who find equivalent widths of $\sim 150$~eV for a flat disk illuminated by X-rays with power-law photon spectral energy distributions. There are at least four ways in which equivalent widths might be elevated above the values we find: (1) super-solar Fe abundance in the disk material, possibly arising as a result of an elevated dust-to-gas ratio; (2) disk flaring, resulting in a solid angle coverage $> 2\pi$; (3) line-of-sight obscuration of the central flaring source (but not fluorescent line photons) by optically thick structures such as the star itself (ie the flare occurring on the far hemisphere); and (4) fluorescence contributions from ionisation by non-thermal electrons. By analogy with the solar case, in which excitation by non-thermal electrons is usually negligible \citep{Parmar.etal:84,Emslie.etal:86} and is much more difficult on energetic grounds, we consider (4) the {\em least} plausible of these. \citet{Ballantyne.Fabian:03} have also shown in the accretion disk context that Fe~K production by non-thermal electron bombardment requires 2-4 orders of magnitude greater energy dissipation in the electron beam than is required for an X-ray photoionization source. Disk flaring can give rise to increased line strengths by factors $< 2$ simply from consideration of the increased solid angle coverage possible compared with an infinite flat disk. It is also difficult to envisage enhanced Fe abundances in the disk being able to elevate line strengths by more than a factor of a few. Of some interest, then, is the observation of an Fe~K$\alpha$ line equivalent width of $\approx1400$~m\AA\ during the rise phase of a flare on the PMS Orion nebula star forming region object V~1486~Ori by \citet{Czesla.Schmitt:07}. Such an enhancement over an infinite disk value of $\sim 150$~m\AA , even with a large degree of disk flaring, would still require extreme enhancements of the disk Fe abundance by an order of magnitude or more \citep[e.g. \S\ref{s:fesens} and][]{Matt.etal:97} were the line due to photoionisation by the {\em directly observed} continuum. We point out, however, that fluorescent line photons from a PMS disk can still be observed when the X-ray flaring source is located behind the star and obscured from the line-of-sight. In the case of the V~1486~Ori flare, the large observed Fe~K$\alpha$ equivalent width is simply and plausibly associated with a partially obscured flare whose rise phase was not fully observed directly owing to line-of-sight obscuration by the star itself. Such obscured flares will inevitably be the cause of some fraction of observed Fe~K$\alpha$ lines from PMS stellar disks. This explanation is also more consistent than one relying on preferential disk geometry and Fe abundance with the non-detection of Fe~K$\alpha$ from a second less extreme flare whose impulsive phase instead appears to have been quite visible. \subsection{Fluorescence from stellar photospheres} The strong flare in II~Peg observed by {\it Swift} and analysed by \citet{Osten.etal:07} presents another interesting case. Photospheric Fe~K$\alpha$ was clearly detected throughout the event. Equivalent widths for different times in the flare ranged from 18 to 61~eV, with uncertainties in the 20-45\%\ range. The authors favoured a collisional excitation mechanism for the line, arguing that fluorescence would be unlikely to produce an observable feature. This assessment employed a simple analytical formula applicable to optically-thin cases in which only a minor fraction of the incident X-ray flux is subject to photoabsorption or scattering \citep[see e.g.][]{Liedahl:99,Krolik.Kallman:87}. \citet{Osten.etal:07} correctly noted that the path length required to obtain the observed equivalent widths under such an approximation was similar to the $\tau=1$ Compton scattering depth, but discounted fluorescence as a possibility on these grounds. Other than the inapplicability of the optically-thin formula for the photospheric fluorescence case, one reason such an argument is overly pessimistic is that incidence angles on the photosphere range from $\sim 0$--$90^\circ$ for small scale heights and path lengths for escape are less than penetration depths by the factor of the inverse cosine of these angles. We defer a more detailed treatment of the event to future work, but note here that equivalent widths of 50~eV are achieved for flare heights up to $\sim 0.2 R_\star$ or so for the $10^8$~K model in our grid, a temperature similar to the average of the values found for the flare by \citet{Osten.etal:07}. While collisional ionisation cannot be ruled out observationally as the source of the observed Fe fluorescence, it is not a requirement. | 7 | 10 | 0710.0621 |
0710 | 0710.5374_arXiv.txt | While isolated neutron stars (INSs) are among the brightest $\gamma$-ray sources, they are among the faintest ones in the optical, and their study is a challenging task which require the most powerful telescopes. \hst\ has lead neutron star optical astronomy yielding nearly all the identifications achieved since the early 1990s. Here, the major \hst\ contributions in the optical studies of INSs and their relevance for neutron stars' astronomy are reviewed. | \label{sec:1} Before the launch of \hst, optical studies of INSs were the exception. In the first 20 years since the pulsars discovery, only the Crab and Vela pulsars were identified (Cocke 1969; Lasker 1976), while optical pulsations were detected from an unidentified source at the center of SNR B0540-69 in the LMC (Middleditch \& Pennypacker 1985), and only a candidate counterpart was found for the misterious $\gamma$-ray source Geminga (Bignami et al. 1987). This score was expected to be considerably improved by \hst, thanks to its much larger sensitivity with respect to ground based telescopes, and to the sharp spatial resolution of the \wfpc\ as well as to the near-UV view of the ESA's \foc. Unfortunately, the spherical aberration of the \hst\ optics affected the execution of most approved proposals, except for those aimed at the brightest targets. So, in the early 1990s the leadership in the INSs optical astronomy was still in the hands of ground-based observatories, mainly in those of the ESO \ntt\ which secured the identification of Geminga through the proper motion of its counterpart, a technique soon become the standard one, and the likely identifications of the optical pulsar in SNR B0540-69 and of PSR B0656+14 (see Mignani et al. 2000). However, the refurbishment of \hst\ in SM-1 (Dec. 2003) brought its performance back to the original expectations and gave it a leading role in INSs' optical astronomy, mantained even after the advent of the 10-m class telescopes. Since \hst\ has provided 8 new INSs identifications, against the 2 of the \vlt\ and the \keck\ (see Mignani et al. 2004), boosting the identification rate by a factor 4. This could have been higher if not for the \foc\ removal in SM-3B (March 2002) and for the \stis\ failure (Aug. 2004), which alone have yielded nearly all the \hst\ INSs identifications, depriving the telescope of its near-UV view. Thus, \hst\ observations have opened wide a new, important observing window on INSs and triggered the interest of a larger and larger fraction of the neutron star community. | 7 | 10 | 0710.5374 |
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0710 | 0710.2248_arXiv.txt | {} {We intend to establish the X-ray properties of Swift J0732.5--1331 and therefore confirm its status as an intermediate polar.} {We analysed 36\,240~s of X-ray data from {\em RXTE}. Frequency analysis was used to constrain temporal variations and spectral analysis used to characterise the emission and absorption properties.} {The X-ray spin period is confirmed to be 512.4(3)~s with a strong first harmonic. No modulation is detected at the candidate orbital period of 5.6~h, but a coherent modulation is present at the candidate 11.3~h period. The spectrum is consistent with a 37~keV bremsstrahlung continuum with an iron line at 6.4~keV absorbed by an equivalent hydrogen column density of around $10^{22}$ atoms~cm$^{-2}$.} {Swift J0732--1331 is confirmed to be an intermediate polar.} | Intermediate polars (IPs) are a sub-class of cataclysmic variables (CVs). They fill the phase space, in terms of magnetic field strength, and spin and orbital periods, between non-magnetic CVs and the strongly magnetic synchronously rotating polars. The magnetic field strength is believed to be in the range of a few MG to tens of MG at the white dwarf surface. This is large enough to dramatically alter the accretion flow, yet not large enough to synchronize the spin and orbital periods. This magnetic field gives rise to the defining characteristic of the sub-class, that of X-ray variation pulsed at the spin period of the white dwarf. For an exhaustive review of CVs see e.g. \cite{warner95}. There are between twenty six\footnote{http://asd.gsfc.nasa.gov/Koji.Mukai/iphome/iphome.html as of 23/8/7.} and fifty IPs currently known (depending on the selection criteria used). The hard X-ray selected object, Swift~J0732.5--1331 (hereafter J0732), is a suspected IP in need of confirmation. The circumstance of its discovery makes J0732 similar to the host of candidate IPs that have been discovered to be powerful emitters of hard X-rays/soft gamma-rays in the 20--100~keV range in the {\sl INTEGRAL\/}/IBIS survey \citep{barlow06}. We have embarked on a campaign of pointed {\sl RXTE\/} observations of these hard X-ray discovered candidate IPs. Here we present the first results of our campaign on J0732. | The unambiguous X-ray spin period detection at 512.4(3)~s, along with the spectral fit to an absorbed 37~keV bremsstrahlung model with an iron line, confirm the intermediate polar status of Swift J0732.5--1331. We are unable to determine the orbital period from these {\em RXTE} data although there is some indication of modulation at the previously suggested photometric period of 11.3~h and none at the spectroscopic period of 5.6~h. To conclude we note that this system is similar, in terms of its small $P_{\rm spin}/P_{\rm orb}$ value, to the IPs RX~J2133.7+5107 and NY~Lup (IGR~J15479--4529). Both of these are {\em INTEGRAL} hard X-ray sources and both also have soft X-ray components. We might therefore expect that Swift~J0732.5--1331 would also display such characteristics upon further study. | 7 | 10 | 0710.2248 |
0710 | 0710.0767_arXiv.txt | We investigate the possibility that near future observations of ultra-high-energy cosmic rays (UHECRs) can unveil their local source distribution, which reflects the observed local structures if their origins are astrophysical objects. In order to discuss this possibility, we calculate the arrival distribution of UHE protons taking into account their propagation process in intergalactic space i.e. energy losses and deflections by extragalactic magnetic field (EGMF). For a realistic simulation, we construct and adopt a model of a structured EGMF and UHECR source distribution, which reproduce the local structures actually observed around the Milky Way. The arrival distribution is compared statistically to their source distribution using correlation coefficient. We specially find that UHECRs above $10^{19.8}$eV are best indicators to decipher their source distribution within 100 Mpc, and detection of about 500 events on all the sky allows us to unveil the local structure of UHE universe for plausible EGMF strength and the source number density. This number of events can be detected by five years observation by Pierre Auger Observatory. | \label{intro} The origin of ultra-high-energy cosmic rays (UHECRs) above $10^{19}$eV is one of the most intriguing problems in astroparticle physics. Akeno Giant Air Shower Array (AGASA) found statistically significant small-scale clusterings of observed UHECR events with large-scale isotropy \citep*{takeda99}. The AGASA data set of 57 events above $4 \times 10^{19}$eV contains four doublets and one triplet within separation angle of $2^{\circ}.5$, consistent with the experimental angular resolution. The chance probability of observing such multiplets under an isotropic distribution is only about 1\% \citep*{hayashida00}. A combination of the results of many UHECR experiments (including AGASA) also revealed eight doublets and two triplets within $4^{\circ}$ on a totally 92 events above $4 \times 10^{19}$eV \citep*{uchihori00}. These multiplets suggest that the origins of UHECRs are point-like sources. For identification of UHECR sources, arrival directions of UHECRs have been observed in detail by High Resolution Fly's Eye (HiRes) and Pierre Auger Observatory (Auger). However, so far, these experiments have reported no significant clustering on the arrival distribution above $4 \times 10^{19}$eV \citep*{abbasi05,mollerach07}. Recently, several classes of astrophysical objects in many literature has tested for positional correlations with observed arrival directions of UHECRs. The correlations with BL Lac objects were discussed on the assumptions of smaller deflection angles of UHECRs than the experimental angular resolution and/or neutral primaries \citep*{tinyakov01}, and in consideration with the deflection by Galactic magnetic field (GMF) \citep*{tinyakov02}. \cite{gorbunov05} considered various classes of powerful extragalactic sources for the correlation study including small corrections of UHECR arrival directions by GMF. \cite{hague07} discussed the correlation with nearby active galactic nuclei (AGNs) from RXTE catalog of AGNs. However, these studies have not taken into account UHECR propagation in extragalactic space. UHECRs above $8 \times 10^{19}$eV lose a significant fraction of their energies by photopion production in collision with the cosmic microwave background (CMB) photons during their propagation \citep*{berezinsky88,yoshida93}. Thus, UHECRs have {\it horizons}, which are the maximum distances of their sources that UHECRs can reach the Earth, even if their energies are below $8 \times 10^{19}$eV at the Earth. The positional correlations between arrival directions of UHECRs and their source candidates outside the horizons are not significant. (In \cite{hague07}, only nearby AGNs within the horizons are considered.) In addition to the UHECR horizons, deflections due to extragalactic magnetic field (EGMF) are also important since extragalactic cosmic rays are propagated for a much greater distance than in Galactic space. Propagation process of UHECRs should be considered in such correlation studies. \cite{yoshiguchi03} investigated the correlation between the arrival distribution of UHECRs and their source distribution taking into account UHECR propagation in intergalactic space with a uniform turbulent EGMF whose strength is 1 nG and coherent length is 1 Mpc. The authors adopted a source distribution with $10^{-6}~{\rm Mpc}^{-3}$ that reproduced the local structures and the AGASA results. They concluded that detection of a few thousand events above $4 \times 10^{19}$eV reveal observable correlation with the sources within 100 Mpc. However, a uniform turbulent field is not realistic EGMF model. Faraday rotation measurements indicate magnetic field strengths at the $\mu$G level within inner region ($\sim$ central Mpc) of galaxy clusters \citep*{kronberg94}. The evidence for synchrotron emission in numerous galaxy clusters \citep*{giovannini00} and in a few cases of filaments \citep*{kim89,bagchi02} also seems to suggest the presence of magnetic fields with $0.1-1.0\mu$G at cosmological structures. Several numerical simulations of large-scale structure formation have confirmed these magnetic structures \citep*{sigl03,dolag05}. Based on these studies, in recent years, we calculated propagation of UHE protons in a structured EGMF which well reproduces the local structures actually observed and simulated their arrival distributions with several normalizations of EGMF strength and several number density of UHECR sources \citep*{takami07}. We constrained the source number density to best reproduce the AGASA results. As a result, $10^{-5}~{\rm Mpc}^{-3}$ is the most appropriate number density, which is weakly dependent on EGMF strength. (In rectilinear propagation, similar number density is also obtained in \cite{blasi04,kachelriess05}) However, this has large uncertainty due to the small number of observed events at present. $10^{-4}~{\rm Mpc}^{-3}$ and $10^{-6}~{\rm Mpc}^{-3}$ are also statistically allowed. Therefore, it is useful to deliberate the correlation between the arrival distribution and the source distribution in the case of these number densities. Note that we revealed in the paper that this uncertainty will be solved by future increase of detected events. In this study, we calculate the arrival distribution of UHECRs, taking their propagation process into account, and investigate the correlation the arrival distribution and their source distribution in the future. A structured EGMF model and source distribution which can reproduce the local universe actually observed are adopted. The source number density and the EGMF strength are treated as parameters since these have some uncertainty. A goal of this study is that we understand the number of observed events to start to observe the UHECR source distribution by UHECRs and how much the correlation is expected in the future. Auger has already detected more events above $10^{19}$eV than those observed by AGASA \citep*{roth07}. Nevertheless, the event clustering has not observed, as mentioned above. It might be due to EGMF and/or GMF strong enough not to generate the multiplets or statistical fluctuation for small number of observed events at highest energies. In any case, we should predict and discuss how the arrival distribution reflects UHECR source distribution. Chemical composition of UHECRs is very important for the correlation. If UHECRs are heavier components, magnetic deflections are larger and the correlation is worse. One of observables for study of UHECR composition is the depth of shower maximum, $X_{\rm max}$, which can be measured by fluorescence detectors. Its average value $\left< X_{\rm max} \right>$ is dependent on UHECR composition and energy. HiRes reported that composition of cosmic rays above $10^{19}$eV is dominated by protons as a result of $X_{\rm max}$ measurement \citep*{abbasi05b}. Recent result by Auger is compatible to the HiRes result within systematic uncertainties \citep{unger07}. However, they concluded that the interpretation of $\left< X_{\rm max} \right>$ distribution is ambiguous because of the uncertainties of hadronic interaction at highest energies. Thus, UHECR composition at highest energies is controversial at present. Despite that, in this study, we assume that all UHECRs are protons since composition above $10^{19}$eV has proton-like feature. This paper is organized as follows: in section \ref{model} we provide our models of UHECR source distribution and a structured EGMF. In section \ref{method} we explain our calculation method for the arrival distribution with UHECR propagation and statistical method. In section \ref{results}, The results of the correlation between the arrival distribution of UHE protons and their source distribution. We summarizes this study in section \ref{conclusion}. | \label{conclusion} In this paper, we calculated the arrival distribution of UHE protons taking into account energy losses and deflections by EGMF during propagation in intergalactic space in order to investigate the possibility that future observations of UHECRs can unveil the local structure of UHE universe. In order to reproduce a realistic situation, we adopted a structured EGMF model and source distributions which reproduce the observed local structures. The arrival distribution of UHE protons was compared statistically to their source distribution using the correlation coefficients. As the number of observed events increases, the correlation coefficient increases and converges to some value which represents the ability to unveil the source distribution by UHE protons, i.e. charged particles. Thus, the number of events that the correlation coefficient starts to converge is an important number for UHECR observations. In other words, detection of such number of events allows us to unravel UHECR source distribution. We found that UHECRs above $10^{19.8}$eV are best indicators to decipher their source distribution within 100 Mpc from discussion based on the final values of the correlation coefficients and GZK mechanism, and 5000, 500, and 200 event detections above $10^{19.8}$eV on all the sky can unveil their source distribution for the source number densities of $10^{-4}$, $10^{-5}$, and $10^{-6}~{\rm Mpc}^{-3}$ respectively. Note that ground based detectors observe only about half hemisphere, so only half of such numbers are requested. In this study, we took only EGMF into account as magnetic field, i.e., neglected GMF GMF deflects trajectories of UHE protons efficiently by its regular components, which consist in spiral and dipole components \citep*{alvarez02,yoshiguchi03b}. A turbulent component of GMF very weakly change the arrival directions of UHE protons \citep*{yoshiguchi04}. The deflection angles of UHECR protons are a few degree at around $10^{20}$eV except for the direction of Galactic Center. Such deflection disturbs the spatial pattern of UHECR arrival distribution at a few degree scale. The effect of GMF is one of our future investigations. A lot of inquiries on UHECR source number density result in around $10^{-5}~{\rm Mpc}^{-3}$ on ground of the AGASA results as introduced in section \ref{intro}. If this number density is true, five year observation by Auger and future observation by TA and Extreme Universe Space Observatory \citep*{euso} will reveal the distribution of nearby UHECR sources. The dawn of the UHE particle astronomy is just around the corner. | 7 | 10 | 0710.0767 |
0710 | 0710.1557_arXiv.txt | \noindent The field of astroparticle physics is currently developing rapidly, since new experiments challenge our understanding of the investigated processes. Three messengers can be used to extract information on the properties of astrophysical sources: photons, charged Cosmic Rays and neutrinos. This review focuses on high-energy neutrinos ($E_{\nu}>100$~GeV) with the main topics as follows. \begin{itemize} \item The production mechanism of high-energy neutrinos in astrophysical shocks. The connection between the observed photon spectra and charged Cosmic Rays is described and the source properties as they are known from photon observations and from charged Cosmic Rays are presented. \item High-energy neutrino detection. Current detection methods are described and the status of the next generation neutrino telescopes are reviewed. In particular, water and ice Cherenkov detectors as well as radio measurements in ice and with balloon experiments are presented. In addition, future perspectives for optical, radio and acoustic detection of neutrinos are reviewed. \item Sources of neutrino emission. The main source classes are reviewed, i.e.~galactic sources, Active Galactic Nuclei, starburst galaxies and Gamma Ray Bursts. The interaction of high energy protons with the cosmic microwave background implies the production of neutrinos, referred to as GZK neutrinos. \item Implications of neutrino flux limits. Recent limits given by the AMANDA experiment and their implications regarding the physics of the sources are presented. \end{itemize} | 7 | 10 | 0710.1557 |
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0710 | 0710.2232_arXiv.txt | We have detected 523 sources in a survey of the Small Magellanic Cloud (SMC) Wing with {\it Chandra}. By cross-correlating the X-ray data with optical and near-infrared catalogues we have found 300 matches. Using a technique that combines X-ray colours and X-ray to optical flux ratios we have been able to assign preliminary classifications to 265 of the objects. Our identifications include four pulsars, one high-mass X-ray binary (HMXB) candidate, 34 stars and 185 active galactic nuclei (AGNs). In addition, we have classified 32 sources as 'hard' AGNs which are likely absorbed by local gas and dust, and nine 'soft' AGNs whose nature is still unclear. Considering the abundance of HMXBs discovered so far in the Bar of the SMC the number that we have detected in the Wing is low. | Multi-wavelength studies of the Small Magellanic Cloud (SMC) have shown that it contains a large number of X-ray binary pulsars. From analysis of H$\alpha$ measurements \citep{ken91} and supernova birth rates \citep{fil98} the star formation rate (SFR) for the SMC is estimated to lie in the range 0.04--0.4 $M_{\odot}$ yr$^{-1}$. \citet{sht05} used these upper and lower SFR estimates and the linear relation between the number of high-mass X-ray binaries (HMXBs) and the SFR of the host galaxy from \citet{gri03} to predict the number of HMXBs expected in the SMC with luminosities $\ge 10^{35}$ erg s$^{-1}$. They found that between 6 and 49 of these systems should be present. Currently $\sim60$ known or probable HMXBs have been detected in the SMC \citep[see e.g.][]{hab04,coe05,mcg07}. It is believed that the considerable number of pulsars can be explained in terms of a dramatic phase of star formation, probably related to the most recent closest approach of the SMC and the Large Magellanic Cloud \citep[LMC;][]{gar96}. To date most of the X-ray studies of the SMC have concentrated on the Bar which has proved to be a significant source of HMXBs. These systems not only provide an homogeneous sample for study, but also give direct insights into the history of our neighbouring galaxy as they are tracers of star formation rates. Part of the puzzle of the X-ray population of the SMC is the missing or under represented components. In particular, there are no known low-mass X-ray binaries (LMXBs) or black hole binaries and only one confirmed supergiant X-ray binary detected to date \citep[see also][]{mcb07a}. A survey of the X-ray binary population of the LMC by \citet{neg02} revealed a similar distribution (within small number statistics) to that in our galaxy - all types were present. It is therefore important to try and identify the ``missing'' X-ray binary types in the SMC. We recently completed the first X-ray survey in the SMC Wing with \chan\ (see Section \ref{sect:obs} for more details). A study of the brightest ($>50$ counts) X-ray sources uncovered two new pulsars, and detected two previously known pulsars \citep{mcg07}. In addition to the four pulsars, the sample included two foreground stars, 12 probable AGNs and five unclassified sources. We found that the pulsars had harder spectra than the other bright X-ray sources. In this paper we report on the analysis of the whole survey and present preliminary classifications for a large fraction of the sources detected. \begin{figure} \includegraphics[width=84mm]{fig1.eps} \caption{The location of the 20 fields studied by Chandra in this work, overlaid on a neutral hydrogen density image of the SMC \citep{sta99}. The Wing and Bar of the SMC are marked.} \label{fig:smc_fields} \end{figure} | \label{sect:disc} For the 523 sources detected in the SMC Wing survey we have been able to find optical matches for 300 of them, and assign preliminary classifications to 265 objects. Our classification method has the advantage that it does not require optical spectra, however, it still requires optical counterparts to be identified. We also note that to classify the remaining 49\% of the survey deeper optical surveys are needed, and in some cases better coverage of the Wing. The majority of the Wing sources are found to be AGNs. In the whole survey we only identify four pulsars \citep[see][]{mcg07} and one HMXB candidate, which compared to the Bar is a small sample. The relatively few pulsars detected in the Wing is perhaps not surprising given the accepted link between regions of H$\alpha$ and star formation, with the main regions of star formation coinciding with the high density H$\alpha$ region in the Bar \citep{ken95}. However, in general, the pulsars we detected in the Wing have harder spectra than those in the Bar. It is also remarkable that the only supergiant system so far detected in the SMC, SMC X-1, lies in the Wing. We note that, despite appearances, the SMC is a very three dimensional object. Studies of the Cepheid population by \citet{lan86} have revealed that the depth of the SMC is up to 10 times its observed width. The two main structures, the Bar and the Wing, could be separated by 10--20 kpc. Could different populations be represented in the two regions? In the case of the HMXBs if we based our response on the X-ray results alone we could perhaps draw the conclusion that the sources in the Wing and Bar are in fact different. However, taking into account the optical spectral analysis in which the optical counterparts for the pulsars were found to be typical of other HMXBs in the SMC \citep{sch07,mcb07b}, different populations seem less likely. This could imply that there is absorption local to the sources which effects the X-ray spectral results. There is also the possibility that a greater population of HMXBs does exist in the Wing of the SMC, but we were not fortunate enough to catch more than a handful of them when they were switched on. From our studies of 10 years of {\it RXTE} data we find that the probability of a Be X-ray transient being in an active phase is only, on average, $\sim 10$\% \citep[Figure 4.62,][]{gal06}. Quiescent X-ray transients have been detected previously in the Milky Way with luminosities $<10^{34}$ erg s$^{-1}$ \citep[e.g.][]{neg00,cam02}. The origin of the quiescent luminosity in Be X-ray transients is still under debate, with a number of processes suggested to account for the detected emission \citep[see e.g.][]{cam02,kre04}. The two mechanisms detectable from sources located in the SMC are: accretion onto the magnetospheric boundary, the propeller regime \citep{ill75,cam00}, and very low rate accretion onto the surface of the neutron star, i.e. residual/leaking accretion \citep[e.g.][]{ste94}. The one HMXB candidate that we have identified has a luminosity (at the distance to the SMC) of $3.2\times10^{33}$ erg s$^{-1}$ so it could be a quiescent source. The lack of HMXBs in the Wing indicates that we are looking at an older population which is confirmed by optical studies of the star formation history of the SMC \citep[e.g.][]{har04}. In theory this should increase our chances of detecting LMXBs. Arguably, LMXBs should be well distributed within the SMC, i.e. they should lie in the Bar and the Wing, however, deep looks of the SMC Bar \citep{naz03} have been unsuccessful in detecting any. The number of LMXBs expected in the SMC is proportional to the total stellar mass of the galaxy, resulting in a prediction of only one system with an X-ray luminosity of $\geq 10^{35}$ erg s$^{-1}$ \citep[see][]{sht05}. However, \citet{gar01} have shown that quiescent LMXBs can be as faint as $2\times 10^{30}$ erg s$^{-1}$. To go as deep as that is beyond the capability of current X-ray telescopes, but in 100 ks it would be possible to reach a limit of $\sim 10^{32}$ erg s$^{-1}$, sufficient to detect a sample of fainter sources and study their characteristics. If an observation like this were performed in the Wing it could be compared directly with the deep exposures of the Bar \citep{naz03,zez05} and help quantify the LMXB population in the SMC. | 7 | 10 | 0710.2232 |
0710 | 0710.0371_arXiv.txt | One well-known way to constrain the hydrogen neutral fraction, $\bxhi$, of the high-redshift intergalactic medium (IGM) is through the shape of the red damping wing of the \lya absorption line. We examine this method's effectiveness in light of recent models showing that the IGM neutral fraction is highly inhomogeneous on large scales during reionization. Using both analytic models and ``semi-numeric" simulations, we show that the ``picket-fence" absorption typical in reionization models introduces both scatter and a systematic bias to the measurement of $\bxhi$. In particular, we show that simple fits to the damping wing tend to \emph{overestimate} the true neutral fraction in a partially ionized universe, with a fractional error of $\sim 30\%$ near the middle of reionization. This bias is generic to any inhomogeneous model. However, the bias is reduced and can even underestimate $\bxhi$ if the observational sample only probes a subset of the entire halo population, such as quasars with large HII regions. We also find that the damping wing absorption profile is generally steeper than one would naively expect in a homogeneously ionized universe. The profile steepens and the sightline-to-sightline scatter increases as reionization progresses. Of course, the bias and scatter also depend on $\bxhi$ and so can, at least in principle, be used to constrain it. Damping wing constraints \emph{must} therefore be interpreted by comparison to theoretical models of inhomogeneous reionization. | \label{intro} The reionization of hydrogen in the intergalactic medium (IGM) is a landmark event in the early history of structure formation, because it defines the moment at which galaxies (and black holes) affected every baryon in the Universe. As such, it has received a great deal of attention -- both observationally and theoretically -- in the past several years. Unfortunately, the existing observational evidence is enigmatic (see \citealt{fan06-review} for a recent review). Electron scattering of cosmic microwave background photons implies that reionization occurred at $z \sim 10$, albeit with a large uncertainty \citep{page06}. On the other hand, \lya forest spectra of quasars at $z \sim 6$ show some evidence for a rapid transition in the globally-averaged neutral fraction, $\bxhi$ (e.g., \citealt{fan06}). However the \lya absorption is so saturated in the \citet{gunn65} trough (with optical depth $\tau_{\rm GP} \ga 10^5 \bxhi$) that constraints derived from that spectral region \citep{fan06, maselli07} are difficult to interpret (e.g, \citealt{lidz06, becker07}). Another probe is the red damping wing of the IGM \lya absorption: the line is so saturated at these redshifts that even photons that are emitted redward of the \lya resonance can suffer significant absorption from the strong damping wings of that transition. This has a number of consequences for high-redshift observations. For example, surveys that search for high-$z$ galaxies through their \lya emission lines will find fewer and fewer galaxies as the IGM becomes more and more neutral \citep{haiman02-lya, santos04}, although galaxy clustering strongly moderates this decline \citep{furl04-lya, furl06-lya, mcquinn07, mesinger07-lya}. Such surveys have now detected objects at $z \sim 6.5$--$9$ (e.g., \citealt{kashikawa06, iye06, stark07}), but their implications for reionization are unclear \citep{malhotra04, haiman05-lya, malhotra06, kashikawa06, dawson07, dijkstra07, mcquinn07-lya, mesinger07-lya}. The evolution of galaxy abundances and clustering measures the damping wing absorption in a statistical sense, but even more information can potentially be gleaned from the damping wing absorption profiles in individual objects \citep{miralda98}. For the galaxies described above, this information is difficult to extract because of their faintness and the complicated origins of their \lya emission lines \citep{mcquinn07-lya}. However, high signal-to-noise spectra of bright objects could be extremely helpful. If the damping wing profile from IGM absorption can be isolated from these spectra, this would provide detailed information on the neutral gas along each particular line of sight (LOS) -- rather than the statistical information available from most other probes. This is very useful, as reionization is expected to be highly inhomogeneous. There are two candidates for such high signal-to-noise spectra at high-redshifts: quasars and gamma-ray bursts (GRBs). Quasars present several challenges: complicated intrinsic spectra, biased IGM environments \citep{barkana04-grb, lidz07}, and large HII regions (which significantly weaken the damping wing absorption redward of the quasar \lya line, and can necessitate detailed spectral analysis of the blue side of the line; \citealt{madau00, mesinger04-mockprox}). Nevertheless, there have already been two claims of damping wing detections in high-redshift spectra, both using quasars from the Sloan Digital Sky Survey (SDSS). \citet{mesinger04} detected a $\bxhi \gsim 0.2$ damping wing through the decreased fluctuations in the total Ly$\alpha$ optical depth near the edge of the HII region surrounding \qnametwoeight\ ($z_S=6.28$). Similarly, by simulating the optical depth distributions blueward of the Ly$\alpha$ line center and comparing them with deep observations, \citet{mesinger07-prox} detected the presence of a $\bxhi \gsim 0.033$ damping wing in the spectra of \qnametwoeight\ and \qnametwotwo\ ($z_S=6.22$). The maximum likelihood was at $\bxhi=1$ for both quasars. The second set of candidates, GRBs, have fewer obstacles to overcome. Long-duration GRBs are believed to be remnants of massive stars (and so trace the bulk of the star formation, which probably occurs in lower-mass halos with more ``typical" IGM environments), and their afterglows have extremely simple power-law intrinsic spectra (see, e.g., \citealt{piran05} for a review). The event rates at high redshifts may be quite high, and cosmological time-dilation helps to identify the sources when they are still bright \citep{bromm02-grb, ciardi00, lamb00, mesinger05}. As a result, there is a great deal of optimism in the literature regarding their potential for damping-wing measurements (e.g. \citealt{miralda98, barkana04-grb}). The highest-redshift GRB afterglow observed so far (at $z \approx 6.3$), has already been used to constrain the global neutral fraction at that time \citep{kawai06, totani06}. Unfortunately, this object illustrates the major difficulty with the red damping wing test for GRBs: intrinsic absorption in the host galaxies \citep{miralda98}. Most GRBs are now known to have large columns of associated neutral hydrogen \citep{vreeswijk04, chen04-grb}. Roughly $20\%$ of well-studied objects have $N_{\rm HI} \la 10^{20} \colden$ \citep{chen07}, although nearly all of the objects in this sample are at $z \la 6$. The $z \approx 6.3$ GRB does appear to have intrinsic absorption with $N_{\rm HI} \sim 10^{21.6} \colden$ \citep{totani06}, which makes it difficult to constrain the IGM absorption. In principle, it is still possible because isolated HI absorbers have different spectral profiles than the IGM (with the optical depth inversely proportional to the wavelength offset squared for isolated absorbers, and to the wavelength offset itself for the IGM). The two sources can then be separated by looking at the shape of the absorption. \citet{totani06} found a best fit with $\bxhi=0$ and estimated that $\bxhi \la 0.17$ ($0.60$) at 68\% (95\%) confidence. Better constraints will require faster followup (when the afterglow is brighter) and systems with less intrinsic absorption. To date, the red damping wing test has generally been assumed to be simple and straightforward. It is usually argued that the absorption is sensitive to a large path length in the IGM, so that small-scale clumpiness can be ignored and that the ionized fraction can be taken to be uniform (for an exception, see \citealt{barkana02}). However, most models of reionization have much more inhomogeneous distributions of neutral and ionized gas, with discrete HII regions surrounding clusters of galaxies, and a sea of nearly neutral gas separating them (e.g., \citealt{arons72, shapiro87}). Such a picture is inevitable when hot stars ionize the gas. Moreover, the most recent models show that the ionized bubbles can become quite large even relatively early in reionization, with sizes $\ga 10 \Mpc$ when $\bxhi \sim 0.5$ \citep{furl04-bub, furl05-charsize, iliev05-sim, zahn07-comp, mcquinn07, mesinger07}. Because the damping wing is sensitive to fluctuations on Mpc scales, it is actually not a good approximation to take the IGM ionized fraction to be constant. In this paper, we will examine whether (and how) the damping wing can actually be used to constrain the reionization process. We summarize the basic physics of the line in \S \ref{lya}. We then examine a series of toy models of the ``picket-fence" absorption typical of the IGM during reionization in \S \ref{toy}. In particular, we show that interpreting measurements with the naive view of a uniform IGM is not only subject to significant scatter (from the different networks of ionized bubbles intersected along different lines of sight) but also a substantial systematic bias. In \S \ref{sim}, we describe the ``semi-numeric" simulations used to generate our main results, which we present in \S \ref{results}. This more detailed picture confirms that scatter between different lines of sight and bias relative to the naive view will be critical in interpreting any observed sources. Finally, we conclude in \S \ref{disc}. When this project was nearing completion, we learned of a similar effort by \citet{mcquinn07-damp} and refer the reader there for a complementary discussion. In our numerical calculations, we assume a cosmology with $\Omega_m=0.26$, $\Omega_\Lambda=0.74$, $\Omega_b=0.044$, $H=100 h \hunits$ (with $h=0.74$), $n=0.95$, and $\sigma_8=0.8$, consistent with the most recent measurements \citep{spergel06}. Unless otherwise specified, we use comoving units for all distances. | \label{disc} In this paper, we have examined how the shape of the \lya red damping wing can be used to constrain the IGM before reionization is complete. In the past, it has usually been assumed that the absorbing gas can be well-approximated by a uniform density medium with constant ionized fraction. However, recent reionization models have shown that ionized bubbles can be quite large, so the latter is not a good approximation. We have therefore critically examined how well the damping wing constrains the neutral fraction during inhomogeneous reionization. We have identified two major issues with its interpretation. First, there is substantial scatter in the optical depth along different lines of sight. Most of this is due to the scatter in the distance between the source and the nearest patch of neutral gas; however, there is still non-negligible scatter even if this distance can be measured from the shape of the damping wing. In our semi-numeric simulations, the fractional r.m.s. fluctuation in $\bxhi$ thus estimated increases from 0.1 to 1 over the range $0.9\gsim\bxhi\gsim0.2$. Fortunately, this statistical uncertainty can be reduced simply by finding more lines of sight. The other problem is more severe: we have shown that the ``picket-fence" absorption from inhomogeneous reionization adds a systematic, and often large, \emph{bias} to measurements of the neutral fraction. Although the damping wing is indeed sensitive to a large path length through the IGM, it is most sensitive to the closest gas. As a result, simple fits to the damping wing will always \emph{overestimate} the true neutral fraction in a partially ionized universe, with an error of $\sim 30\%$ near the middle of reionization. This bias is generic to any inhomogeneous model. The bias is reduced and can even become negative if observations only probe a subset of the entire halo population, such as quasars with large HII regions. Both the systematic and statistical uncertainty can be reduced by a careful fit to the damping wing spectral profile, which is typically steeper than the naively expected $(\Delta \lobs)^{-1}$ profile. However, because the absorption typically comes from many neutral patches, a large number of parameters are required for a detailed fit, and given the relatively modest difference from the expected behavior, these will be difficult to measure, probably only possible in systems with intrinsically large optical depths. Moreover, the scatter in the profiles, even at fixed $\tau_D$, is sufficient that large samples will be required to put strong constraints on reionization from the spectral shape. Of course, the bias and scatter also depend on $\bxhi$ and so can, at least in principle, be used to constrain it. For example, large dispersion in the inferred neutral fractions could be an indicator of $\bxhi \la 0.2$. If an independent estimate of $\bxhi$ exists, one could reverse the direction of analysis, and use the bias and scatter to constrain the reionization model and topology. Fortunately, for a given model of reionization, the dispersion and bias can be calibrated by theoretical models. We therefore argue that the most efficient way to constrain reionization with the damping wing is through comparison with detailed models. Of course, any such constraints will be model-dependent, but we believe that the morphology of reionization is now sufficiently well-understood (see, e.g., \citealt{furl05-charsize, mcquinn07}) that these uncertainties will likely not dominate the statistical uncertainties from the small number of accessible sources, at least in the relatively near future. For example, the reionization morphology is nearly independent of redshift \citep{furl04-bub, mcquinn07}. Also, we have found only a modest dependence of the $x_D$ distribution on halo mass (mostly due to the variation in bubble size with mass). However, toward the end of reionization, when the absorption is dominated by rare, narrow sheets of neutral hydrogen, the details of the radiative transfer algorithm (or an approximation to it, as in our models) and of the sample selection will be extremely important. Nevertheless, the task is challenging, as the damping wing profile must be separated from the rapidly varying resonance absorption for quasars (as in \citealt{mesinger04, mesinger07-prox}) or from intrinsic absorbers for GRBs. Fortunately, in the latter case $\sim 20\%$ of moderate-redshift GRBs have only modest absorbers and will still be useful for these purposes \citep{chen07}. So far, the damping wing analysis has been performed on three high-redshift quasars: \qnamefourtwo\ ($z_S=6.42$), \qnametwoeight\ ($z_S=6.28$), \qnametwotwo\ ($z_S=6.22$) \citep{mesinger04, mesinger07-prox}, as well as GRB 050904 ($z_S \approx 6.3$) \citep{kawai06, totani06}. This paper highlights the need to calibrate these and future damping wing analysis with simulations of the reionization morphology. Obviously we cannot set firm constraints without detailed simulations of the observations. Nevertheless, the mean bias we find from our simulations seems to work in the direction of strengthening the upper limit (on $\bxhi$) from the \citet{totani06} measurements, and weakening the lower limit from the \citet{mesinger04, mesinger07-prox} constraints at $\bxhi\lsim0.6$ (although, interestingly, it would strengthen them if $\bxhi\gsim0.6$; see the sign change for the bias in Fig.~\ref{fig:high_mass_mean_sig}). Conversely, the steeper-than-expected absorption profile seems to work in the direction of weakening the \citet{totani06} constraints (especially because it must be distinguished from strong internal absorption) while strengthening the \citet{mesinger04, mesinger07-prox} constraints. The absorption profile might be more relevant than the bias for these studies, as an overall bias can be partially degenerate with other free parameters in the fit: that is, when the absorption profile \emph{can} be detected to high precision, its shape will certainly be useful in constraining $\bxhi$. The scatter in both effects would probably somewhat erode the confidence contours for all of these studies. On the other hand, our model predicts large scatter between different LOSs at the end of reionization, which is consistent with the measurements at $z \sim 6.3$. More precise limits will require simultaneous fits to the intrinsic absorption and the range of possible IGM absorber profiles, and we defer them to future work. Another intriguing possibility is to try to measure damping wing characteristics from stacked spectra of many \lyans-emitting galaxies. \citet{mcquinn07} have shown that the wing shape is difficult to separate from uncertainties in the line for individual objects, and the scatter we have described will also make the interpretation of individual faint emitters problematic. But, if the characteristics of the population are relatively constant, stacking may increase the signal to noise sufficiently to allow a detection of a ``mean" damping wing at each redshift, even far redward of line center. SRF thanks Crystal Martin and Josh Bloom for conversations that stimulated this work. We thank Z. Haiman for helpful comments on this manuscript. This research was partially supported by grant NSF-AST-0607470. | 7 | 10 | 0710.0371 |
0710 | 0710.2697_arXiv.txt | We studied the clustering properties and multiwavelength spectral energy distributions of a complete sample of 162 Ly$\alpha$-emitting (LAE) galaxies at $z\simeq 3.1$ discovered in deep narrow-band MUSYC imaging of the Extended Chandra Deep Field South. LAEs were selected to have observed frame equivalent widths $>$80\AA\ and emission line fluxes $>1.5\times10^{-17}$ergs cm$^{-2}$ s$^{-1}$. Only 1\% of our LAE sample appears to host AGN. The LAEs exhibit a moderate spatial correlation length of $r_0=3.6^{+0.8}_{-1.0}$Mpc, corresponding to a bias factor $b=1.7^{+0.3}_{-0.4}$, which implies median dark matter halo masses of $\log_{10}\mathrm{M_{med}} = 10.9^{+0.5}_{-0.9}$M$_\odot$. Comparing the number density of LAEs, $1.5\pm0.3\times10^{-3}$Mpc$^{-3}$, with the number density of these halos finds a mean halo occupation $\sim$1--10\%. The evolution of galaxy bias with redshift implies that most $z=3.1$ LAEs evolve into present-day galaxies with $L<2.5L^*$, whereas other $z>3$ galaxy populations typically evolve into more massive galaxies. Halo merger trees show that $z=0$ descendants occupy halos with a wide range of masses, with a median descendant mass close to that of $L^*$. Only 30\% of LAEs have sufficient stellar mass ($>\sim3\times10^9$M$_\odot$) to yield detections in deep Spitzer-IRAC imaging. A two-population SED fit to the stacked $UBVRIzJK$+[3.6,4.5,5.6,8.0]$\mu$m fluxes of the IRAC-undetected objects finds that the typical LAE has low stellar mass ($1.0^{+0.6}_{-0.4}\times10^9$M$_\odot$), moderate star formation rate ($2\pm1 $M$_\odot$yr$^{-1}$), a young component age of $20^{+30}_{-10}$Myr, and little dust ($A_V<0.2$). The best fit model has 20\% of the mass in the young stellar component, but models without evolved stars are also allowed. | \label{sec:intro} The discovery of high-redshift Ly$\alpha$-emitting galaxies (LAEs) opened a new frontier in astronomy \citep{cowieh98,huetal98}. Because the Ly$\alpha$ line is easily quenched, a galaxy with detectable Ly$\alpha$ emission is likely dust-free, i.e., in the initial phases of a burst of star formation. The Ly$\alpha$ lines have large equivalent widths ($20${\AA}$<$EW$_{\mathrm{rest}}<\sim 100${\AA}) and broad velocity widths (100 km~s$^{-1}$$<$FWHM$<$800 km~s$^{-1}$) and are often asymmetric, indicative of high-redshift galaxies undergoing active star-formation \citep[e.g.,][]{manningetal00, kudritzkietal00,arnaboldietal02,rhoadsetal03, dawsonetal04,huetal04,venemansetal05,matsudaetal06,gronwalletal07}. Other high-redshift galaxy populations (including Lyman break galaxies, Distant Red Galaxies, Sub-Millimeter Galaxies) exhibit strong clustering and should evolve into elliptical and giant elliptical galaxies today \citep{adelbergeretal05a,quadrietal07a,webbetal03}. These objects were selected by unusually strong rest-frame continuum emission in the ultraviolet, optical, and far-IR respectively, resulting in $10^{10}L_\odot \leq L_{bol} \leq 10^{12}L_\odot$ \citep{reddyetal06b}. Such strong continua appear to occur primarily in deep potential wells that are strongly biased versus the general distribution of dark matter halos. LAEs instead offer the chance to probe the faint end of the (bolometric) luminosity function of high-redshift galaxies, which contains the majority of galaxies. The strong Ly$\alpha$ emission line allows detection and spectroscopic confirmation of LAEs with typical bolometric luminosities $\simeq 10^{10} L_\odot$. A detailed calculation of the LAE luminosity function at $z\simeq 3.1$ is given in \citet{gronwalletal07}. Spectral energy distribution (SED) modelling of the stacked $UBVRIzJK$ photometry of 18 LAEs in the Extended Chandra Deep Field-South (ECDF-S) \citep{gawiseretal06b} showed the average galaxy to have low stellar mass ($<10^9$ M$_\odot$) and minimal dust \citep[see also][]{nilssonetal07}. LAEs have the highest {\it specific} star formation rates (defined as SFR divided by stellar mass) of any type of galaxy, implying the youngest ages \citep{castroceronetal06}. Because Ly$\alpha$ emission is easily quenched by dust, LAEs have often been characterized as protogalaxies experiencing their first burst of star formation \citep[e.g.][]{hum96}. However, the differing behavior of Ly$\alpha$ and continuum photons encountering dust and neutral gas makes it possible for older galaxies to exhibit strong Ly$\alpha$ emission when morphology and kinematics favor the escape of these photons towards Earth \citep{neufeld91,haimans99,hanseno06}. This could allow older, dusty galaxies with actively star-forming regions to exhibit Ly$\alpha$ emission with high equivalent width. SED modelling of LAEs using Spitzer-IRAC \citep{fazioetal04} to probe rest-frame near-infrared wavelengths, where old stars dominate the emission, has yielded mixed results. \citet{pirzkaletal07} report extremely young ages of a few Myr and low stellar masses ($10^6$M$_\odot<$M$_*<10^8$M$_\odot$) from SED modelling of 9 LAEs at $4.0<z<5.7$ in the Hubble Ultra Deep Field. However, \citet{laietal07a} performed SED fitting to 3 LAEs with IRAC detections out of a sample of 12 $z=5.7$ LAEs in GOODS-N and found ages as high as 700 Myr and significant stellar masses ($10^9<$ M$_\odot$ $<10^{10}$), making it appear that these LAEs were not undergoing their first burst of star formation. The 9/12 LAEs lacking IRAC detections are presumably younger and less massive and might lack an evolved population. Investigating the nature of the LAEs without IRAC detections requires a stacking approach to see if the typical LAE stellar mass is low enough to have been generated in a single ongoing starburst. Stacking will yield the best results when applied to a large statistical sample of LAEs in a region with deep IRAC imaging. \citet{gronwalletal07} present the largest available sample of LAEs in an unbiased field, 162 LAEs at $z=3.1$ in the ECDF-S discovered as part of the MUSYC survey (\citealp{gawiseretal06a}, \url{http://www.astro.yale.edu/MUSYC}). The ECDF-S has been targeted with deep narrow-band imaging and multi-object spectroscopy, complemented by public broad-band $UBVRIzJK$, Spitzer+IRAC and Chandra+ACIS-I imaging. We improve the constraints of \citet{gawiseretal06b} using the MUSYC $UBVRIzJK$ photometry of our larger sample of LAEs and adding IRAC [3.6$\mu$m,4.5$\mu$m,5.8$\mu$m,8.0$\mu$m] Cycle 2 legacy images from SIMPLE (Spitzer IRAC/MUSYC Public Legacy of the ECDF-S, \url{http://www.astro.yale.edu/dokkum/simple}). This Letter summarizes our imaging and spectroscopic observations of LAEs, presents our results from clustering analysis and SED modelling, and discusses the implications for the formation process of typical present-day galaxies. We assume a $\Lambda$CDM cosmology consistent with WMAP results \citep{spergeletal07} with $\Omega_m=0.3, \Omega_\Lambda=0.7$, $H_0 = 70$ km s$^{-1}$ Mpc$^{-1}$, and rms dark matter fluctuations on $8h^{-1}$ Mpc scales given by $\sigma_8 = 0.9$ . All correlation lengths and number densities are comoving. We have suppressed factors of $h_{70}$ in reporting correlation lengths, number densities, dark matter masses, stellar masses and star formation rates. | \label{sec:discussion} In CDM cosmology, galaxy formation is an ongoing process caused by merging of lower-mass dark matter halos, which may already possess stars. Finding stellar population ages of $<100$ Myr is interesting. Our analysis of halo merger trees from the Milli-Millenium simulation \citep{springeletal05} found the median age of dark matter halos with M$>10^{10.6}$M$_\odot$ at $z=3.1$ (defined as the age since half of the dark matter mass was accumulated) to be $\sim$600 Myr, with only $<$10\% of halos younger than 100 Myr. If repeated LAE phases occur, the mean halo occupation of $\sim1$--$10$\% can be interpreted as a ``duty cycle'' telling us what fraction of each halo's lifetime is spent in the early phases of starbursts before significant dust is generated, and the population-averaged young age of $\sim$20 Myr would imply that this phase typically lasts $\sim$40 Myr. Alternatively, if all dark matter halos experience a single LAE phase shortly after their ``formation'' in a major merger, the mean halo occupation implies that LAEs will only be found in the youngest $1$--$10$\% of halos, which is barely consistent with their single-population best-fit age of 100 Myr. If LAEs represent a subset of dark matter halos selected to have ages less than 100 Myr, their observed clustering may underestimate their dark matter halo masses by up to a factor of two \citep[see][]{gaow07}. \begin{figure}[h!] \epsscale{1.25} \plotone{f6.eps} \caption[] { Tracks show the evolution of bias with redshift calculated using the no-merging model. The filled circle shows our result for the bias of LAEs at $z=3.1$. Previous results at high-redshift are shown for LAEs at $z=4.5$ and $z=4.86$ (stars, \citealp{kovacetal07} and \citealp{ouchietal03}, respectively), LBGs at $z\sim4$ (stars, \citealp{ouchietal04b}), K-selected galaxies (diamonds, \citealp{quadrietal07a}), bright LBGs at $z\sim 3$ (triangle, \citealp{leeetal06}), and BM, BX, and LBG galaxies (asterisks, \citealp{adelbergeretal05b}, \citealp{adelbergeretal05a}). Local galaxy clustering is shown for SDSS galaxies (open circles, \citealp{zehavietal05}) and for rich clusters (cross, \citealp{bahcalletal03}). K-band limits are in Vega magnitudes. } \label{fig:bias} \end{figure} Figure \ref{fig:bias} shows the reported bias values for LAEs to be lower than those of other $z>3$ galaxy populations \citep[bias values determined as in][]{quadrietal07a}. The expected evolution of bias is shown for the ``no-merging'' model \citep{fry96,whiteetal07}. A realistic amount of merging will cause the bias to drop somewhat faster, so the plotted trajectories provide an upper limit on the bias factor of a given point at lower redshifts. This shows that typical $z=3.1$ LAEs will evolve into galaxies of at most a few times $L^*$ at $z=0$. The bias values imply that LAEs at $z=3.1$ might evolve into the subset of BX galaxies at $z\simeq 2.2$ dimmer than $K=21.5$, which also show relatively weak clustering \citep{adelbergeretal05a}. The $K>21.5$ BX galaxies have average M$_*=1.5\times10^{10}$M$_\odot$, so the $z=3.1$ LAEs would need to form stars at an average rate of 14 M$_\odot$yr$^{-1}$ over the intervening Gyr. This could be achieved with a constant specific SFR and no merging or with a constant SFR and $\sim$2 major mergers. The only previous measurements of LAE clustering in unbiased fields are at $z=4.5$ \citep{kovacetal07} and $z=4.86$ \citep{ouchietal03}, and this earlier LAE population appears to have significantly stronger clustering, consistent with possibly evolving into typical Lyman break galaxies at $z\simeq 3$. The models of \citet{ledelliouetal06} predict stellar and dark matter masses and star formation rates for $z=3$ LAEs within a factor of two of our results, despite assuming a top-heavy IMF and a very low escape fraction $f_{esc}=0.02$ that appears inconsistent with the observed lack of dust \citep[see][for an alternative approach]{kobayashietal07}. \citet{maoetal07} used the stellar mass of $5\times10^8$M$_\odot$ observed by \citet{gawiseretal06b} to predict LAE dark matter masses of $10^{10}<$M$<10^{11}$M$_\odot$, in the lower end of our allowed range. The stellar ages of $\sim 20$ Myr preferred by the two-population fit are noticeably lower than the maximum values of 100 to 500 Myr predicted by these authors, \citet{moriu06}, and \citet{haimans99}, but the ages of 60 to 350 Myr preferred for the case of no evolved stars would be compatible. None of the current models and numerical simulations of LAEs \citep[see also][]{thommesm05,razoumovs06,tasitsiomi06} predict their present-day descendants. Nonetheless, the evolution of a significant fraction of $z=3.1$ LAEs into $z=0$ $L^*$ galaxies with dark matter mass $M_{DM}\simeq2\times10^{12}$M$_\odot$ and stellar mass $M_*\simeq4\times10^{10}$M$_\odot$ \citep{ichikawaetal07} appears reasonable. Fig. \ref{fig:mass} shows the histogram of present-day masses of dark matter halos in the Milli-Millenium merger trees that have progenitors with M$>5\times10^{10}$M$_\odot$ at $z=3.1$. The median present-day halo mass is $1.2\times10^{12}$M$_\odot$, and this would increase if LAEs found in sub-halos of massive $z=3.1$ halos were included. \citet{lietal07} predict that the main progenitor of a present-day $L^*$ galaxy had a dark matter mass of $\sim10^{11}$M$_\odot$ at $z=3$ and that these galaxies experienced several major mergers at $1.5<z<7$. To form an $L^*$ galaxy at $z=0$, several LAEs could merge while experiencing a mild reduction in average SFR, with accretion of lower-mass dark matter halos through minor mergers providing most of the final dark matter mass. However, the halo mass distribution of $z=0$ descendants in Fig.\ref{fig:mass} is very broad, with 25th and 75th percentile values of $2.9\times10^{11}$M$_\odot$ and $7.6\times10^{12}$M$_\odot$. While $z=0$ $L^*$ galaxies like the Milky Way are roughly the median descendants of $z=3.1$ LAEs, the descendant halos include a wide range from dwarf galaxies to rich galaxy groups. \begin{figure}[h!] \epsscale{1.25} \plotone{f7.eps} \caption[] { Histogram of dark matter halo masses of present-day descendants of halos with M$>5\times10^{10}$M$_\odot$ at $z=3.1$. The dashed lines show the median halo mass of $1.2\times10^{12}$M$_\odot$ and the 25th and 75th percentile values of $2.9\times10^{11}$M$_\odot$ and $7.6\times10^{12}$M$_\odot$. } \label{fig:mass} \end{figure} The typical LAE stellar mass at $z=3.1$ is lower than that of any other studied high-redshift population \citep[see][]{reddyetal06b} but is close to that of dim ($i<26.3$) Lyman break galaxies (LBGs) at $z\sim5$ \citep{vermaetal07}. LAEs at $z=3.1$ have much lower star formation rate, stellar age, stellar mass, dark matter halo mass, and dust extinction than the $\sim30$M$_\odot$yr$^{-1}$, $\sim500$ Myr, $\sim2\times10^{10}$M$_\odot$, $\sim3\times10^{11}$M$_\odot$, $A_V\simeq1$ LBG population at $z\sim3$ \citep[$R<25.5$,][]{shapleyetal01,adelbergeretal05a} or the $\sim100$M$_\odot$yr$^{-1}$, $\sim2$ Gyr, $\sim10^{11}$M$_\odot$, $\sim10^{13}$M$_\odot$, $A_V\simeq2.5$ Distant Red Galaxy (DRG) population \citep{webbetal06,forsterschreiberetal04,quadrietal07a}. The high-redshift Sub-Millimeter Galaxies \citep{chapmanetal03b} appear to be the most massive and dusty, with the highest SFR. LAEs may represent the beginning of an evolutionary sequence where galaxies gradually become more massive and dusty due to mergers and star formation, but most LAEs at $z=3.1$ will never reach the DRG stage since DRG stellar and dark matter masses are already greater than those of present-day $L^*$ galaxies. The Damped Ly$\alpha$ Absorption systems (DLAs, \citealp{wolfeetal05}) are another high-redshift population that probes the faint end of the luminosity function. The dark matter halo masses of DLAs at $z\sim3$ were determined by \citet{cookeetal06b} to lie in the range $10^9<$M$<10^{12}$M$_\odot$ i.e., $1.3<b<4$, which overlaps with the range of both $L^*$ and super-$L^*$ progenitors in Fig. \ref{fig:bias}. At least half of the DLAs appear to have ongoing star formation \citep{wolfeetal04} and two of the three DLAs detected in emission were seen in Ly$\alpha$. Further study is needed to determine the relationship between DLAs and LAEs. The observed properties of LAEs at $z=3.1$ make them the most promising candidates to be high-redshift progenitors of present-day $L^*$ galaxies like the Milky Way. Our results suggest that LAEs are observed during the early phases of a burst of star formation, perhaps caused by a major merger of smaller dark matter halos. The input halos appear to have already contained stars, accounting for the evolved stellar population that appears to contribute most of the LAE stellar mass, although starburst-only models are also allowed. It is clear that not all progenitors of L$^*$ galaxies were LAEs at $z=3.1$. The comoving number density of our sample of LAEs is a factor of 15 less than $\phi^*$ for local galaxies \citep{linetal96}, plus we expect several high-redshift halos to merge into a single galaxy today. It remains possible that all progenitors of present-day galaxies experienced an LAE phase at {\it some} redshift. Clustering and SED studies of LAEs at various redshifts are needed to assess the validity of this hypothesis. | 7 | 10 | 0710.2697 |
0710 | 0710.1694_arXiv.txt | An eight stage balanced modulation scheme for dual beam polarimetry is presented in this paper. The four Stokes parameters are weighted equally in all the eight stages of modulation resulting in total polarimetric efficiency of unity. The gain table error inherent in dual beam system is reduced by using the well known beam swapping technique. The wavelength dependent polarimetric efficiencies of Stokes parameters due to the chromatic nature of the waveplates are presented. The proposed modulation scheme produces better Stokes $Q$ and $V$ efficiencies for wavelengths larger than the design wavelength whereas Stokes $U$ has better efficiency in the shorter wavelength region. Calibration of the polarimeter installed as a backend instrument of the Kodaikanal Tower Telescope is presented. It is found through computer simulation that a $14\%$ sky transparency variation during calibration of the polarimeter can introduce $\approx 1.8\%$ uncertainty in the determination of its response matrix. | \label{sec:intro} Polarimetric accuracy is one of the most important goals in modern astronomy. It is limited since most optical elements encountered by the light on its path from the source to the detector, can alter its state of polarization (for eg. telescope optics, imaging system, grating, etc). Apart from these, variation in sky transparency, image motion and blurring due to the atmosphere are a major concern in high precision ground based solar polarimetry. The effect of atmosphere, which is commonly known as seeing induced effect, can be reduced by fast modulation schemes(Stenflo and Povel, 1985). The modulation frequencies in these schemes are generally higher than seeing fluctuations, which is $\approx 1kHz$(Stenflo and Povel, 1985 and Lites, 1987). Large format CCDs, which are required to cover reasonable spectral and spatial range, will pose difficulty in reading out the data at $kHz$ speed. Stenflo and Povel(1985) proposed a scheme whereby rapidly modulated signal is demodulated by optical means, thereby avoiding the need to read the detectors at a rapid rate. Lites(1987) has proposed a rotating waveplate modulation scheme as an alternative to minimize the seeing induced cross-talk among Stokes parameters. There he has shown that the faster the rotation rate of the modulator, the lower the cross-talk among Stokes parameters. And the seeing induced cross-talk levels of a dual beam system are factors 3-5 smaller than those of a single beam system. However, in dual beam system, the error introduced due to flat field residual is a matter of concern in high precision polarimetry. A possible solution to the above mentioned problems can be found by using a mixed scheme in which spatial and temporal modulations are performed(Elmore et al. 1992, Martinez Pillett et al. 1999 and Sankarasubramanian et al. 2003). The gain table uncertainties are avoided using the beam swapping technique(Donati et al. 1990, Semel et al. 1993 and Bianda et al. 1998) A low cost dual beam polarimeter has been installed as a backend instrument for the Kodaikanal Tower Telescope (KTT). Different modulation schemes were studied and an optimum scheme is identified. The proposed scheme requires eight stages of modulation of input light in order to obtain the maximum polarimetric efficiency. Laboratory experiments have been performed to verify the theoretical understanding of the proposed scheme. The studies are extended to other wavelengths apart from the design wavelength of $\lambda6300$. The outline of this paper is as follows. The proposed eight stage modulation scheme for the measurement of general state of polarization is discussed in section(2). Wavelength dependence of the efficiency of the polarimeter in measuring Stokes parameters is presented in section(3). In section(4), the performance of the polarimeter at KTT is presented. | An eight stage modulation scheme to measure the general state of polarization is presented here. Beam swapping technique is incorporated in this scheme, which helps in alleviating the gain correction errors. The total polarimetric efficiency is close to unity as the Stokes parameters are weighted equally in all the stages of modulation. The final Stokes parameters are demodulated using all the stages of intensity measurements. Hence, the derived input Stokes parameters are equally weighted time averaged quantities over the time of measurement. Since the retarders used in the polarimeter are chromatic, the efficiency of the polarimeter in measuring Stokes $QUV$ is wavelength dependent. The laboratory experiments performed to study the wavelength dependence of efficiency of the polarimeter confirms the theoretical expectations. It is found through computer simulation that a $14\%$ sky transparency variation can cause $\approx 1.8\%$ uncertainty in the elements of the polarimetric response matrix during its calibration, for the modulation/demodulation scheme presented here. The non-zero values of the off-diagonal elements are not a serious concern if those values do not change drastically in short time scales. During any solar polarimetric measurements, data for the calibration of the polarimeter are taken at least once a day. Calibration of the polarimeter are carried out on a few days over a period of 10-day and the response matrix derived over this period did not show any appreciable variations. The variations in the off-diagonal elements are less than the fit errors ($<$ 0.5\%). The polarization signals observed on the Sun is always less than 40\% and hence an uncertainty of 0.5\% in the calibration will produce an inaccuracy of 0.2\% in the polarization signals. The measured total polarimetric efficiency of the polarimeter installed at KTT is $\approx 0.986$ at $\lambda6563$ wavelength region which is better than some of the polarimeters such as ZIMPOL(0.72), ASP(0.88), TIP(0.92) and POLIS(0.84). | 7 | 10 | 0710.1694 |
0710 | 0710.3691_arXiv.txt | We developed a source detection algorithm based on the Minimal Spanning Tree (MST), that is a graph-theoretical method useful for finding clusters in a given set of points. This algorithm is applied to $\gamma$-ray bidimensional images where the points correspond to the arrival direction of photons, and the possible sources are associated with the regions where they clusterize. Some filters to select these clusters and to reduce the spurious detections are introduced. An empirical study of the statistical properties of MST on random fields is carried in order to derive some criteria to estimate the best filter values. We introduce also two parameters useful to verify the goodness of candidate sources. To show how the MST algorithm works in the practice, we present an application to an EGRET observation of the Virgo field, at high galactic latitude and with a low and rather uniform background, in which several sources are detected. | Telescopes for satellite-based high energy $\gamma$-ray astronomy detect individual photons by means of the electron-positron pair that they generate through the detector. From the pair trajectories it is possible to reconstruct the original direction of the photon with an uncertainty that decreases with the energy, from a few degrees below 100 MeV to less than a degree above 1 GeV. This technique was applied to the past $\gamma$-ray observatories SAS-2 (Fichtel et al. 1975), COS-B (Bennett 1990) and EGRET-CGRO (Kanbach et~al. 1988; Thompson et~al. 1993), all equipped with spark chambers. Pair tracking is also used in the current AGILE mission (Tavani et al. 2006) and in the LAT telescope on board the next GLAST mission, both employing silicon microstrip detectors (Gehrels et al. 1999). The resulting product is an image where each photon is associated with a direction in the sky: discrete sources thus correspond to regions in which a number of photons higher than those found in the surroundings are observed. When the size of this region is consistent with the instrumental Point Spread Function the source is considered as point-like, otherwise it can be extended or a group of near sources. Various algorithms are applied to the detection of point-like or extended sources in $\gamma$-ray astronomy: the most extensively used one is based on the Maximum Likelihood (Mattox et al. 1996), whereas others based on Wavelet Transform analysis (Damiani et al. 1997), Optimal Filter (Sanz et al. 2001), Scale-Adaptive Filter (Herranz et al. 2002), etc., were variously applied to real and simulated data to study their performances. Some of them are based on deconvolution techniques of the instrumental Point Spread Function (PSF). Many methods work directly on the pixellated images, i.e. count or intensity maps. Other methods search for clusters in the arrival directions of photon that, if statistically significant, are considered an indication of a source. The approach considered by us is essentially a cluster search based on a \emph{minimal spanning tree} (MST) algorithm. This technique has its root in graph theory, and highlights the \emph{topological} pattern of connectedness of the detected photons. Given a graph $G(V, E)$, where V is the set of vertices (or \emph{nodes}) and E is the set of weighted \emph{edges} connecting them, a MST (Kruskal 1956; Prim 1957; Zahn 1971) is the tree (a subgraph of $G$ without closed circuits) that connects all the points with the minimum total weight, defined as the sum of the weight of each tree's edge. In a data set consisting of points in a Cartesian frame of reference, we can consider them as the nodes of a graph, the edges being the lines joining the nodes, weighted by their length. The MST method was originally proposed for $\gamma$-ray source detection by Di~Ges\`u and Sacco (1983), who investigated also the statistical properties in uniform fields. This work was developed by Di~Ges\`u and Maccarone (1986), and De~Biase et al. (1986) applied MST for detecting extended sources in EXOSAT X-ray images. Other authors applied MST methods to the goal of finding galaxy clusters, both in 2 and 3-dimensional surveys and simulations (Barrow et al. 1985; Bhavsar \& Ling 1988a,b; Plionis et al. 1992; Krzevina \& Saslaw 1996, Doroshkevich et al. 2001, 2004) and showed the capabilites of the method as a filament-finding algorithm. In this paper we investigate the MST approach in $\gamma$-ray source detection, and present a new study of its statistical properties and the definition of selection criteria. We also introduce some parameters useful to classify the reliabilty of detected clusters to be associated with source candidates. We would like to emphasize here that this method is not \emph{alternative} to other source detection algorithms, but it is \emph{complementary}, in the sense that it can give a list of possible candidate sources (identified via their photons' clusterization properties) that could be further investigated by other means. This paper is structured as follows. In Sect.~2 we describe our MST algorithm, and in Sect.~3 and 4 we investigate by means of numerical simulations the statistical distributions of edge length and node number, and we introduce some criteria useful for the source detection with our method. An example of application to an EGRET field is shown in Sect.~5, while in Sect.~6 we summarize and discuss our results. | We presented an application of a Minimal Spanning Tree algorithm to the problem of source detection in $\gamma$-ray images. This method does not involves in the computation the instrumental response functions and works recognizing the regions of the sky where arrival directions of photons clusterize. It has the advantages of a fast calculation but did not provide directly estimates of the source flux. We have shown that a MST based algorithm is a viable method to detect $\gamma$-ray sources both in simulated images and in real $\gamma$-ray observations of the EGRET experiment on board Compton-GRO. We proposed some tools to optimize the filtering parameters and to assess the reliability of source detections, like the clustering degree and the bootstrap detection stability. These tools are based on a study, although empirical, of the statistical properties of the Minimal Spanning Tree on random fields. The MST application to an EGRET field around the two famous $\gamma$-ray loud quasars 3C~273 and 3C~279 found almost all the 3EG sources already detected in the same pointing and confirmed the presence of another source, detected in a different pointing. We consider this result a good indication that MST method is particularly efficient. We found also evidence of a new possible source with a significance comparable to that of other well established sources. We expect that future experiments with a better sensivity, like the LAT instrument on board GLAST, will confirm or disprove this finding. There are, however, several possible effects that make difficult the source detection and require even more attention when the MST method is used. These problems can be divided into four main categories: $i$) problems due to the presence of strong sources, $ii$) problems arising from energy spectra of the sources different from that of the background; moreover, different spectral indices between the sources will result in different probabilities to be detected, due to the energy dependence of the PSF, $iii$) problems originated by images with a non-homogeneous background, $iv$) problems due to the geometrical distortions from the arriving celestial photons in projection onto the $\gamma$-ray telescope, that will result not necessarily in a circular shape to characterize proper cluster selections. At present we have not developed a well established strategy to solve these problems and in the following we will briefly discuss some aspects useful for the understanding of results. One or more strong sources in the field have various possible consequences. A first relevant effect is that they are characterized by a high clustering degree and consequently reduce the value of $\Lambda_{\mathrm{m}}$ with respect to the one expected in the field if they were absent. A value of $\Lambda_{\mathrm{c}}$ very close to $\Lambda_{\mathrm{m}}$ would here be good to detect strong sources but this selection criterion could miss other possible sources of lower flux. Another effect is the presence of possible ``satellites'' in the surroundings of a strong source, even closer than expected from the PSF, originated by cutting an edge whose length is just smaller than $\Lambda_{\mathrm{c}}$. For example, the cluster detected in the EGRET field (see Sect.~4) with no obvious counterpart, is at a distance of about 3.3$^{\circ}$ from the strong radio quasar 3C~279, and therefore we cannot exclude that it could be a satellite of the latter. Usually, the satellites do not have a high frequency in the bootstrap fields. The energy distribution of the photons also affects the source detection, because the PSF of $\gamma$-ray telescopes changes with the energy becoming much narrower at high energies. This implies that sources with spectra harder than the background are better detected in high energy images because their clustering degree increases. At variance, sources with soft spectra give more disperse clusters and cannot be easily found. Another class of problems is present when the background is markedly non-homogeneous, as in the case where the field contain a portion of the galactic disc. In this case, using an unique $\Lambda_{\mathrm{c}}$ in all the image would correspond to a long cutting in the dense region and to a short cutting in the region of low density with the consequence of missing real sources and producing more spurious clusters. A general approach to be used for $\gamma$-ray source detection is that of using several methods, possibly based on different techniques, and to compare their results. In this way it will be possible to reduce the number of spurious detections, because of the different criteria and \emph{a priori} assumptions applied in the source recognition. Accordingly, MST method can be used to obtain a quick list of photon clusterization regions, that could correspond to possible sources, to be studied indipendently with other methods. There are other clustering algorithms that can be applied to $\gamma$-ray source detection, like the Voronoi tessellation (Icke \& van de Weygaert 1987, Aurenhammer 1991). In particular, this method is based on the construction of its dual graph, the Delaunay triangulation, of which MST is a subset. We think, therefore, that at least in principle, they would provide similar result and that a combined figure of merit for source detection should be defined. Here we discussed gamma-ray astronomy as a prime candidate for the application of MST method, but it could be even better applicable to the study of data clusterization in ultra-high energy cosmic rays (UHECR) and hemispherical neutrino experiments, that are characterized to the absence of structured background. We think also that it will be possible to extend MST to higher dimensional spaces introducing time and energy as additional dimensions. Basically there are two approaches: $i$) to search for clusters in separate, dimensionally homogeneous subspaces, and then to search for the intersection of the detected clusters and $ii$) to define a new metric for the tree edges that combine together the various dimensions in a suitable way for the MST computation. Preliminary numerical attempts based on the second approach, with energy as third coordinate, seem to be very promising to identify sources having spectra different from that of the background. Another possible 3-dimensional generalization is to take into account also the time, thus searching for variable or stable sources. We will discuss a possible application of such a generalized MST in a subsequent work. | 7 | 10 | 0710.3691 |
0710 | 0710.4552_arXiv.txt | We have analyzed the redshift-dependent fraction of galactic bars over 0.2$<$z$<$0.84 in 2,157 luminous face-on spiral galaxies from the COSMOS 2-square degree field. Our sample is an order of magnitude larger than that used in any previous investigation, and is based on substantially deeper imaging data than that available from earlier wide-area studies of high-redshift galaxy morphology. We find that the fraction of barred spirals declines rapidly with redshift. Whereas in the local Universe about 65\% of luminous spiral galaxies contain bars (SB$+$SAB), at $z\sim0.84$ this fraction drops to about 20\%. Over this redshift range the fraction of {\em strong} (SB) bars drops from about 30\% to under 10\%. It is clear that when the Universe was half its present age, the census of galaxies on the Hubble sequence was fundamentally different from that of the present day. A major clue to understanding this phenomenon has also emerged from our analysis, which shows that the bar fraction in spiral galaxies is a strong function of stellar mass, integrated color and bulge prominence. The bar fraction in very massive, luminous spirals is about constant out to z$\sim$0.84 whereas for the low mass, blue spirals it declines significantly with redshift beyond z=0.3. There is also a slight preference for bars in bulge dominated systems at high redshifts which may be an important clue towards the co-evolution of bars, bulges and black holes. Our results thus have important ramifications for the processes responsible for galactic downsizing, suggesting that massive galaxies matured early in a dynamical sense, and not just as a result of the regulation of their star formation rate. | How, when and at what rate did the Hubble sequence form? This question is central to the field of galaxy formation and evolution. We examine it by measuring the evolution of the bar fraction with redshift using the 2-square degree Cosmic Evolution Survey (COSMOS). In nearly all simulations, the formation timescale for a bar is rapid {\sl once the necessary conditions (a massive, dynamically cold and rotationally-supported disk)} are met. Therefore the redshift evolution of the bar fraction is a fundamental probe of the evolutionary history of disk galaxies. The bar fraction is defined simply as: \begin{equation} f_{\rm bar} = {{\rm number \ of\ barred\ spirals}\over {\rm number\ of\ all\ spirals}}. \end{equation} In the local Universe the value of $f_{bar}$ is quite well established. When only strongly barred\footnote{Bars that are highly elliptical and have rectangular isophotes are classified as strongly barred (SB) galaxies whereas those with more oval shapes are classified as SAB or ovally distorted galaxies} galaxies (SB) are counted, the RSA, RC3 and UGC \citep{sandage87,devau91,nilson73} all give values of $f_{bar}=0.25-0.3$. When ovally distorted (SAB) are also counted the situation becomes a little less clear-cut, because, unlike the RC3, the UGC and RSA do not attempt to carefully compile an inventory of such galaxies. If ovally distorted systems in the RC3 are included in the computation of $f_{bar}$ then the local bar fraction rises to $f_{bar}\sim0.6$. This result is in good agreement with recent infrared studies which have measured the local bar fraction to be $\sim$ 0.65 \citep{eskridge00, whyte02, menendez07, marinova07}. In the infrared, a majority of the SAB galaxies are classified as strongly barred SB systems \citep{eskridge00}. As noted by \citet{eskridge00} and \citet{menendez07}, the overall bar fraction is the same in the infrared and the optical (although there is a small number of cases where bars are unveiled at infrared wavelengths). This is not surprising, because bars are primarily stellar structures whose visibility only declines sharply at ultraviolet wavelengths, short wards of the Balmer break (see also \S \ref{bandshifting}). We conclude that the consensus value of the local barred fraction is $f_{bar}\sim0.3$ for strongly barred systems, and $f_{bar}\sim0.65$ for all barred galaxies, and that these values are so well-known that they have not changed significantly in over four decades. In sharp contrast with the rapid and stable consensus reached on the local bar fraction, attempts to measure the bar fraction at high redshift have proven difficult. The earliest analyses of the bar fraction in the Hubble Deep Fields (HDFs) found a dramatic paucity of barred spirals at z$>$0.5 \citep{abraham96, vanden96, abraham99}. These authors concluded that at lookback times greater than 5 Gyr disks were either dark matter dominated or dynamically too hot (perhaps due to the increased merging activity) to host bars. However, the small volume probed by the HDFs (only thirty bright, face-on spiral galaxies between $0<z<1$) led to concerns that the bar fraction at high redshift may not be adequately measured. \citet{sheth03} investigated whether a significant number of bars could have been missed, as suggested by \citet{bunker99} using the H-band NICMOS HDF. \citet{sheth03} found four bars and two candidate bars out of 95 galaxies at z$>$0.7. Overall, the fraction of barred spirals in the NICMOS HDF remained extremely low, as in the optical HDF studies. But \citet{sheth03} noted that their study was limited by the coarse NICMOS resolution (0$\farcs$15) such that only the largest (and rarest) bars could be identified (bars with semi-major axis $>$5 kpc). When the fraction of these large bars at z$>$0.7 was compared to local samples, there was no compelling evidence for a decline in barred spirals, but likewise the NICMOS data did not unveil any new bars at low redshifts; all except one of the four bars in the \citet{sheth03} study are at z$>$0.9, where k-correction effects are important (\S \ref{bandshifting}). A major advance in spatial resolution was possible with the Advanced Camera for Surveys (ACS) whose 0$\farcs$05 pixels are able to resolve all but the smallest (nuclear, $<$2 kpc diameter) bars at all redshifts. Using ACS data, two studies \citep{elm04,jogee04} found that contrary to the previous HDF results, the bar fraction is constant at 30\% over the last 8 Gyr (since $z=1.2$). The sample sizes, however remained modest in these studies (186 in \citealt{elm04}, and 258 in \citealt{jogee04}). In this paper we examine in detail the redshift evolution of the bar fraction using the unparalleled wide and deep 2-square degree COSMOS data set. The plan for the paper is as follows: in Section 2 we describe our sample selection procedure. The classification methodology we have adopted is described in Section 3. Our main results are presented in Section 4, before being discussed in Section 5. Our conclusions are summarized in Section 6. An Appendix to this paper provides a detailed analysis of possible selection effects at high redshift and a discussion of our local calibration sample of 139 galaxies from the Sloan Digital Sky Survey (SDSS) Data Release 4 \citep{adelman06}. Throughout this paper we adopt a flat $\Lambda$-dominated cosmology with $H_0$=70 km\ s$^{-1}$\ Mpc$^{-1}$, $\Omega_M=0.3$, and $\Omega_\Lambda=0.7$. | Bars are an important signpost of galaxy evolution because once a galaxy disk is sufficiently massive, dynamically cold and rotationally supported it forms a bar. Therefore the evolution of the bar fraction over time is an important indicator of the evolutionary history of disk galaxies and the assembly of the Hubble sequence. Using a detailed analysis of 2,157 L$^*$ face-on, spiral galaxies from 0.0$<$z$<$0.84 in the COSMOS 2-square degree survey we have investigated the evolution of the bar fraction over the last 7 Gyr. We have undertaken an extensive and careful analysis of selection effects (k-correction, surface brightness dimming, inclination, spatial resolution, etc.) which is detailed in the Appendix. Our main results are as follows: $\bullet$ The bar fraction for L$^*$ galaxies drops from about 65\% in the local Universe to about 20\% at z=0.84. Over this redshift range the fraction of {\em strong} bars (SB) drops from about 30\% to under 10\%. Thus at a lookback time of 7 Gyr, when the Universe was half its present age, fundamental aspects of Hubble's `tuning fork' classification sequence had not yet fallen into place. Only about one fifth of all spiral galaxies were ``mature'' enough (dynamically cold, massive and rotationally supported) to host galactic structures of the type we see today. $\bullet$ For the total f$_{bar}$ (SB+SAB), the change is far less dramatic between z=0.3 and z=0.0 indicating slow evolution in galactic structures in L$^*$ galaxies over the last 4 Gyr. It is likely that there is significant evolution in the formation of bars in the sub-L$^*$ galaxies over this period. $\bullet$ One of the most significant findings in this study is the correlation between f$_{bar}$ and the galaxy mass, luminosity and color. We find that in the highest redshift bins f$_{bar}$ is higher in the more massive, luminous and redder systems. In fact, in the most massive systems, f$_{bar}$ is already as high at z=0.8 as the local value. These systems thus had already arrived with their present Hubble types at a lookback time of 7 Gyr. In the subsequent 3 Gyr, from z=0.84 to z=0.3, the lower mass, bluer systems evolved more slowly toward their present Hubble types. Thus the signature of downsizing is intimately connected with dynamical maturity of disks and is present in the formation of galactic structure. $\bullet$ Finally, we find a slight preference for barred galaxies to be more bulge-dominated in the high redshift bin. This correlation is consistent with the dynamical downsizing found for bars in general if bars and bulges both form earlier and more prominently in the most massive galaxies. The lack of a stronger correlation may be related to the variety of bulges: bars are also likely to be involved with the inflow that builds pseudo-bulges. Given the strong correlation between bulge properties and black hole mass seen today, there may be a co-evolution of bars, bulges and black holes in some galaxies. The exact details of these processes remain to be investigated. | 7 | 10 | 0710.4552 |
0710 | 0710.3528_arXiv.txt | {The High Energy Stereoscopic System (H.E.S.S.), located in the Khomas Highlands of Namibia, is an array of four imaging atmospheric-Cherenkov telescopes designed to detect $\gamma$-rays in the very high energy (VHE; $>$ 100 GeV) domain. Its high sensitivity and large field-of-view (5$^{\mbox{\tiny o}}$) make it an ideal instrument to perform a survey within the Galactic plane for new VHE sources. Previous observations in 2004/2005 resulted in numerous detections of VHE gamma-ray emitters in the region l = 330$^{\mbox{\tiny o}}$ - 30$^{\mbox{\tiny o}}$ Galactic longitude. Recently the survey was extended, covering the regions l = 280$^{\mbox{\tiny o}}$ - 330$^{\mbox{\tiny o}}$ and l = 30$^{\mbox{\tiny o}}$ - 60$^{\mbox{\tiny o}}$, leading to the discovery of several previously unknown sources with high statistical significance. The current status of the survey will be presented.} \begin{document} | The majority of the newly discovered sources of very high energy (VHE; $>$ 100 GeV) $\gamma$-rays are related to late phases of stellar evolution, either directly to massive stars or to the compact objects they form after their collapse. The possible associations include pulsar wind nebulae (PWN) of high spin-down luminosity pulsars such as G\,18.0$-$0.7 \cite{hess_j1825}, supernova remnants like RX\,J1713.7$-$3946 \cite{RXJ1713}, and open star clusters like Westerlund\,2 \cite{westerlund2}. As these objects cluster closely along the Galactic plane, a survey of this region is an effective approach to discover new sources and source classes of VHE $\gamma$-ray emission. | The H.E.S.S. Galactic plane survey, which started in the year 2004, now reaches from $-$85$^{\circ}$ longitude to 60$^{\circ}$ longitude, and covers an approximately 6$^{\circ}$ broad band around latitude b = 0$^{\circ}$. In total, more than 950\,hours of data were taken in this region, including survey mode observations, re-observations of source candidates and dedicated observations of known or suspected $\gamma$-ray sources. The first stage of the survey, covering the inner 60$^{\circ}$ of the Galactic plane, has increased the number of known VHE $\gamma$-sources within this region from three at the beginning of 2004 to seventeen. Further follow-up observations within this region and the extension of the survey along the Galactic plane resulted in the discovery of even more additional VHE $\gamma$-ray emitters. Most of them were presented during this conference. Multi-wavelength follow-up observations and archival searches have already begun, and will be crucial for understanding the underlying processes at work in these astrophysical objects. | 7 | 10 | 0710.3528 |
0710 | 0710.1836_arXiv.txt | We report on the first detection of maser emission in the $J$=11-10, $J$=14-13 and $J$=15-14 transitions of the $v$=0 vibrational state of SiS toward the C-rich star IRC+10216. These masers seem to be produced in the very inhomogeneous region between the star and the inner dust formation zone, placed at $\simeq$5-7~R$_*$, with expansion velocities below 10~km~s$^{-1}$. We interpret the pumping mechanism as due to overlaps between $v$=1-0 ro-vibrational lines of SiS and mid-IR lines of C$_2$H$_2$, HCN and their $^{13}$C isotopologues. The large number of overlaps found suggests the existence of strong masers for high-$J$ $v$=0 and $v$=1 SiS transitions, located in the submillimeter range. In addition, it could be possible to find several rotational lines of the SiS isotopologues displaying maser emission. | The detection of strong maser emission at the frequencies of pure rotational transitions of some molecules is a common phenomenon in circumstellar envelopes (CSE's) of evolved stars \citep{elitzur_1992,gray_1999}. The maser is usually produced in a small region of the envelope and sometimes provides valuable information on the physical conditions of the emitting region. Due to the different chemistry, masers are produced by different molecules in O- and C-rich stars. In O-rich stars, SiO exhibits strong maser emission in different rotational transitions within several vibrational states, from $v$=1 to 4 \citep{cernicharo_1993,pardo_1998}. These masers are formed in a region of the CSE very close to the stellar surface and seem to be driven by NIR radiation \citep{pardo_2004}. In C-rich stars, although SiO is present with similar abundances than in O-rich stars \citep{schoier_2006}, no SiO maser has been detected. The explanation could be that SiO is formed at $\simeq$3-5~R$_*$, where the angular dilution of the star is high and the density and temperature lower than in the regions where SiO masers are produced in O-rich stars \citep{agundez_2006}. In C-rich stars only HCN shows strong maser emission in several pure rotational lines within vibrational states from $\nu_2$=1 to 4 \citep{lucas_1989,schilke_2003}. These masers must be formed in the innermost regions of the CSE. SiS has been previously found to show weak maser emission in the $J$=1-0 $v$=0 transition in IRC+10216 \citep{henkel_1983}. In this letter, we report on the first detection of maser emission from the $J$=11-10, 14-13 and 15-14 transitions in the $v$=0 vibrational state of SiS (hereafter M$_1$, M$_2$, and M$_3$) observed toward the C-rich star IRC+10216. We have also obtained observations of $v$=1 rotational lines which exhibit thermal emission. We propose that overlaps of $v$=1-0 ro-vibrational transitions of SiS with mid-IR lines of C$_2$H$_2$ and HCN could provide the pumping mechanism for these masers as well as higher-$J$ $v$=0 SiS masers in the submillimeter range. This discovery is interesting because this species could play in C-rich stars a role similar to that of SiO in O-rich stars: the energy level pattern of both molecules is similar and it is also formed close to the star, as chemical equilibrium and interferometric observations imply \citep{bieging_1989}. | The $v$=1 rotational lines show single cusped profiles and their relative intensities indicate that the rotational levels are thermally populated. The linewidths correspond to velocities of 9-11~km~s$^{-1}$, lower than the terminal velocity $\simeq$14.5~km~s$^{-1}$ \citep{cernicharo_2000}, thus the emission arises from the innermost region of the CSE, between the photosphere and the inner dust formation zone, placed at $\simeq$5~R$_*$ \citep{keady_1993}. We can derive the SiS abundance in that region from the $v$=1 lines assuming an uniform sphere at a distance of 180 pc with a radius of 10 R$_*$, illuminated by the central star (R$_*$=5$\times$10$^{13}$ cm, T$_*$=2300~K), with T$_\textnormal{\scriptsize{k}}$=1000~K and n$_{\textnormal{\scriptsize{H}}_2}$=1.6$\times$10$^{9}$~cm$^{-3}$ (mean values for this region derived by \citealt{fonfria_2006}) and an expansion velocity of 11~km~s$^{-1}$. We solve the statistical equilibrium for SiS considering 100 rotational levels and 3 vibrational states and apply the LVG radiative transfer formalism using the code developed by J. Cernicharo. With the assumed n$_{\textnormal{\scriptsize{H}}_2}$, the column density for H$_2$ is $\simeq$7$\times$10$^{23}$~cm$^{-2}$ and the derived one for SiS is $\simeq$5.0$\times$10$^{18}$~cm$^{-2}$. Hence, the SiS abundance is $\simeq$7$\times$10$^{-6}$. This result is compatible with a higher SiS abundance in the innermost CSE (from LTE chemistry models, 3$\times$10$^{-5}$, \citealt{agundez_2006}) and an abundance further away of 6.5$\times$$10^{-7}$ according to observations of $v$=0 rotational lines over the outer CSE by \citet{bieging_1989}. Most $v$=0 rotational lines show a rounded or slightly double peaked profile with the blue part absorbed by cold SiS through the envelope. However, M$_1$, M$_2$ and M$_3$ show extra emission in the form of narrow peaks (FWHM=1-3~km s$^{-1}$). The velocities of these features, within the $-10$ to 10~km s$^{-1}$ range, indicate that SiS maser emission arises from different regions located between the star and the inner dust formation zone ($r$$\simeq$5~R$_*$). The lines M$_2$ and M$_3$ have been previously observed by \citet{sahai_1984} but no maser emission was noticed. This could be due either to the limited sensitivity of their observations or to a time variability of the SiS maser phenomenon. The bottom-center panel of Fig.~\ref{fig:figure} shows in detail the line profiles of the observed masers. Up to ten maser features labelled ($a$,\ldots,$j$) are identified. The most complex line profile is that of M$_3$. It is formed by 5 main features: $a$, $d$, $f$, $i$, $j$ with $v$=$-8.8$, $-3.7$, $0.88$, $6.3$ and $8.4$~km s$^{-1}$, respectively. The line profiles show that the strongest peaks are at negative velocities, having their red counterparts rather weak. This behavior was previously found by \citet{henkel_1983} for the $v$=0 $J$=1-0 line, which mostly consists of a narrow peak (FWHM=0.3~km~s$^{-1}$) centered at $-13.5$~km~s$^{-1}$. The strong asymmetry of the lineshapes can be either due to blanking by the star of the redshifted maser or by amplification of the blueshifted emission by the foreground stellar environment. The strongest observed maser, M$_3$, with T$_\textnormal{\scriptsize{MB}}$$\simeq$60~K (F$_\textnormal{\scriptsize{obs}}$$\simeq$300~Jy; with thermal and non-thermal emission), is weak compared to some SiO masers detected in O-rich stars \citep*[e.g.][]{cernicharo_1993} or to HCN masers observed in IRC+10216 \citep{lucas_1989,schilke_2003}. However, M$_1$, M$_2$, and M$_3$ are stronger than the SiS $J$=1-0 maser observed towards IRC+10216 by \citet{henkel_1983}. The similarity between maser features in M$_2$ and M$_3$ indicates that they may arise from the same regions and produced by the same pumping mechanism; the maser in M$_1$ is probably formed in other regions. Hence, we suggest two possible geometries of the innermost CSE to explain the observed features: ($i$) An onion-like innermost region, where each maser is produced in a shell. This hypothesis is supported by the symmetry of features $a$--$j$ and $d$--$i$. The peaks at extreme velocities, $a$ and $j$, would be produced just in front of and behind the star near the inner dust formation shell ($r$$\simeq$5~R$_*$) with expansion velocities of $\simeq$5-11~km~s$^{-1}$. The features $d$ and $i$ would be produced in a similar way but in an inner shell with a lower expansion velocity. Finally, the central peak, $f$, would be formed in a shell very close to the star, with the whole shell contributing to the maser emission. M$_1$ would be produced in a cap-shaped region in front of the star. ($ii$) All the masers are formed in different positions of a clumpy shell. The different features in M$_2$ and M$_3$ would be produced in different regions of the shell: peaks $a$ and $j$ in front of and behind the star and the other peaks ($d$, $f$, and $i$) in different clumps, as occur with the only feature of M$_1$. The classic pumping mechanism for the SiO $v$$>$0 masers observed in O-rich stars resides in the increase of the trapping lifetime (A/$\tau$)$_{v \rightarrow v-1}$ with $J$ for $v$$\rightarrow$$v$-1 transitions, when they become optically thick \citep{kwan_1974}. Such mechanism produces masers in adjacent rotational lines of the $v$ state, and explains the $v$=1 and 2 SiO masers \citep{bujarrabal_1981,lockett_1992}. However, the masers observed in rotational transitions of $^{29}$SiO, $^{30}$SiO, and in $v$=3 and 4 of SiO do not show the latter behavior and have been interpreted as due to IR overlaps between ro-vibrational lines of SiO isotopologues \citep{cernicharo_1991,gonzalezalfonso_1997}. For SiS, the absence of maser emission in $v$=1 rotational lines and the odd $v$=0 pattern also exclude the Kwan \& Scoville pumping mechanism. This suggests that overlaps of $v$=1-0 ro-vibrational transitions of SiS with those of mode $\nu_5$ of C$_2$H$_2$ and mode $\nu_2$ of HCN, could provide the pumping mechanism. C$_2$H$_2$ and HCN are abundant in the inner CSE of IRC+10216 and dominate the 11-14 $\mu$m spectrum \citep{fonfria_2006}. Overlaps with these two species have been already proposed by \citet{sahai_1984} to explain the different profiles of adjacent $J$ lines of SiS. However, the SiS frequencies used in that work were not as accurate ($\sigma$$\sim$10$^{-1}$~cm$^{-1}$) as those available today. We have calculated those frequencies from the Dunham coefficients determined by \citet{sanz_2003}, for which the error of the band center is $<$$10^{-4}$~cm$^{-1}$ ($\simeq$0.04~km~s$^{-1}$; the relative accuracy of P and R lines is much better). The frequencies of C$_2$H$_2$, H$^{13}$CCH, HCN, and H$^{13}$CN lines have been taken from the HITRAN Database 2004 \citep{rothman_2005}, with an accuracy better than $10^{-3}$ cm$^{-1}$ ($\simeq$0.4~km~s$^{-1}$) for C$_2$H$_2$ and H$^{13}$CCH, and $10^{-4}$~cm$^{-1}$ for HCN and H$^{13}$CN. Table~\ref{tab:frequencies_sis} shows the mid-IR line overlaps of SiS with C$_2$H$_2$, HCN, and their most abundant isotopologues. For the overlap search we selected coincidences within $|\Delta v|$$<$10~km~s$^{-1}$. However, since the CSE is expanding, every region of the envelope is receding from the others. Hence, if the population of the SiS levels is affected by an overlap with a strong line of other species, the frequency of this overlapping transition must be higher than the SiS one. This would restrict the condition to positive $\Delta v$. Nevertheless, due to the linewidth, lines at $\Delta v$$<$0 can overlap the SiS lines. Hence, we have set the negative cutoff to one half of the typical linewidth in the innermost CSE ($\simeq$5~km~s$^{-1}$; \citealt{fonfria_2006}). Therefore, our search is restricted to $-2.5\le\Delta v\le 10$~km~s$^{-1}$. All the ro-vibrational SiS lines commented hereafter refer to $v$=1$\rightarrow$0 transitions and will be labelled with the usual spectroscopic nomenclature R, Q, P (see footnote of Table~\ref{tab:frequencies_sis}). \begin{deluxetable}{c@{}c|c@{}c@{}c} \tabletypesize{\scriptsize} \tablewidth{0pc} \tablecolumns{5} \tablecaption{Mid-infrared Line Overlaps of SiS with C$_2$H$_2$, HCN, and Their Most Abundant Isotopologues} \tablehead{\colhead{Line} & \colhead{$\nu$ (cm$^{-1}$)} & \colhead{Mol.} & \colhead{Transition} & \colhead{$\Delta v$(km/s)}} \startdata \multicolumn{5}{c}{\textbf{Overlaps Involving SiS Levels of Observed Lines}}\\[2pt] R 9 & 750.3695 & HCN & $01^{1}0-00^{0}0$ R$_e\left(12\right)$ & $-6.6$ \\ P10 & 738.2889 & C$_2$H$_2$ & $1^{-1}1^{1}-1^{-1}0^{0}$ R$_f\left(3\right)$ & \phs{}1.2 \\ R13 & 752.6431 & C$_2$H$_2$ & $0^{0}1^{1}-0^{0}0^{0}$ R$_e\left(9\right)$ & \phs{}5.9 \\[2pt] P14 & 735.7324 & C$_2$H$_2$ & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(38\right)$ & $-6.1$ \\ P15 & 735.0861 & C$_2$H$_2$ & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(36\right)$ & $-9.9$ \\[2pt] \multicolumn{5}{c}{\textbf{Overlaps Involving SiS Levels of Unobserved Lines}}\\[2pt] P 2 & 743.2624 & C$_2$H$_2$ & $0^{0}1^{1}-0^{0}0^{0}$ R$_e\left(5\right)$ & \phs{}0.6 \\ P16 & 734.4368 & C$_2$H$_2$ & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(34\right)$ & \phs{}1.1 \\ P20 & 731.8114 & C$_2$H$_2$ & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(24\right)$ & \phs{}9.5 \\ P22 & 730.4815 & C$_2$H$_2$ & $1^{-1}1^{1}-1^{1}0^{0}$ Q$_e\left(18\right)$ & $-2.1$ \\ R22 & 757.5819 & C$_2$H$_2$ & $1^{1}1^{1}-1^{1}0^{0}$ R$_e\left(10\right)$ & \phs{}0.1 \\ R22 & 757.5819 & H$^{13}$CN & $01^{1}0-00^{0}0$ R$_e\left(17\right)$ & \phs{}5.2 \\ P23 & 729.8123 & H$^{13}$CCH & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(19\right)$ & \phs{}8.7 \\ P24 & 729.1402 & C$_2$H$_2$ & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(1\right)$ & \phs{}9.4 \\ P25 & 728.4654 & H$^{13}$CCH & $0^{0}1^{-1}-0^{0}0^{0}$ Q$_e\left(7\right)$ & $-0.0$ \\ R25 & 759.1733 & HCN & $01^{1}0-00^{0}0$ R$_e\left(15\right)$ & \phs{}3.6 \\ R26 & 759.6977 & C$_2$H$_2$ & $0^{0}1^{1}-0^{0}0^{0}$ R$_e\left(12\right)$ & $-1.0$ \\ R33 & 763.2816 & H$^{13}$CN & $01^{1}0-00^{0}0$ R$_e\left(19\right)$ & \phs{}6.1 \\ P38 & 719.4363 & C$_2$H$_2$ & $1^{1}1^{1}-1^{1}0^{0}$ P$_e\left(5\right)$ & \phs{}4.5 \\ R40 & 766.7130 & C$_2$H$_2$ & $0^{0}1^{1}-0^{0}0^{0}$ R$_e\left(15\right)$ & \phs{}4.3 \\ R42 & 767.6652 & C$_2$H$_2$ & $1^{1}1^{-1}-1^{1}0^{0}$ R$_e\left(20\right)$ & \phs{}8.7 \\ R45 & 769.0697 & C$_2$H$_2$ & $0^{0}1^{1}-0^{0}0^{0}$ R$_e\left(16\right)$ & $-1.8$ \\ P48 & 712.1720 & C$_2$H$_2$ & $1^{-1}1^{-1}-1^{-1}0^{0}$ P$_f\left(8\right)$ & \phs{}4.7 \enddata \tablecomments{Overlaps of SiS with C$_2$H$_2$, H$^{13}$CCH, HCN and H$^{13}$CN found in the mid-IR range $\left[690,780\right]$~cm$^{-1}$ with $-2.5\le\Delta v\le 10$ km~s$^{-1}$, where $\Delta v/c=[\nu (\textnormal{X})-\nu (\textnormal{SiS})]/ \nu (\textnormal{SiS})$ and $J\le 50$, (corresponding SiS $v$=0 rotational frequencies below 900~GHz). The lines and frequencies at the left correspond to the vibrational transitions $v$=1$\rightarrow$0 of SiS. The errors on the velocities are $< 0.5$~km~s$^{-1}$. The notation for the C$_2$H$_2$ and H$^{13}$CCH vibrational states involved in the ro-vibrational transitions is $\nu_4^{\ell_4}\nu_5^{\ell_5}$, whereas for HCN and H$^{13}$CN is $\nu_1\nu_2^\ell\nu_3$. The parity of the lower level is even ($e$) and odd ($f$). The transitions are labelled as R, P, and Q for $J_{up}-J_{low}=+1$, $-1$, 0.} \label{tab:frequencies_sis} \end{deluxetable} In order to qualitatively interpret the effects of IR overlaps on maser emission, we have used the same LVG radiative transfer code modified to account for overlaps, changing the intensity at the overlapping frequency and the escape probability for photons from the overlapped lines. Thus, for some SiS lines, the excitation temperature becomes negative (maser activation) and the brightness temperature is considerably enhanced. M$_2$ and M$_3$ are naturally explained by the overlap of the R$\left(13\right)$ line of SiS with the strong C$_2$H$_2$ line $0^01^1-0^00^0$R$_e\left(9\right)$ (Table~\ref{tab:frequencies_sis}). SiS molecules are easily excited from $v$=0 $J$=13 to $v$=1 $J$=14 through R$\left(13\right)$ and decay to the $v$=0 $J$=15 via P$\left(15\right)$, creating a population inversion between the $v$=0 $J$=13 and 15, and producing maser emission in M$_2$ and M$_3$. M$_1$ may be produced by the overlap of the P$\left(10\right)$ SiS line with the strong C$_2$H$_2$ line $1^{-1}1^1-1^{-1}0^0$R$_f\left(3\right)$. This overlap can pump SiS molecules from $v$=0 $J$=10 to $v$=1 $J$=9 through P$\left(10\right)$, depopulates the $v$=0 $J$=10 level and produces the inversion between $v$=0 $J$=10 and 11. We have also looked for overlaps of $v$=1-0 higher-$J$ SiS lines with C$_2$H$_2$ and HCN transitions to try to predict SiS masers at submillimeter wavelengths. Some of them are shown in the second block of Table~\ref{tab:frequencies_sis}. They suggest, for example, that a maser could be found in rotational transitions involving the $v$=0 $J$=23 level (likely $J$=24-23 and maybe 23-22), or the $v$=0 $J$=25 and 26 states (perhaps in the $v$=0 $J$=27-26 and maybe 26-25). These overlaps could also produce masers in $v$=1 rotational transitions. There are many overlaps of other SiS isotopologues with lines of C$_2$H$_2$, HCN and $^{28}$Si$^{32}$S. With the adopted criteria for the overlap search (see footnote of Table~\ref{tab:frequencies_sis}), we have found 91, 93, 76, and 94 coincidences for $^{29}$SiS, $^{30}$SiS, Si$^{34}$S, and Si$^{33}$S, respectively. Consequently, although these species are less abundant than $^{28}$Si$^{32}$S, the population of some levels could be inverted producing maser emission. This study represents the discovery of three new SiS masers and should be complemented with future observations of higher-$J$ $v$=0 and 1 rotational transitions. Furthermore, a detailed multi-molecule non-local radiative transfer model would help to understand the dust formation region and the role of SiS in C-rich evolved stars. | 7 | 10 | 0710.1836 |
0710 | 0710.0600_arXiv.txt | % We describe the status of a project whose main goal is to detect variability along the extreme horizontal branch of the globular cluster NGC~6752. Based on Magellan 6.5m data, preliminary light curves are presented for some candidate variables. By combining our time-series data, we also produce a deep CMD of unprecedented quality for the cluster which reveals a remarkable lack of main sequence binaries, possibly pointing to a low primordial binary fraction. | Among field B-type subdwarf (sdB) stars, three types of non-radial pulsators have so far been detected, namely: \begin{itemize} \item EC 14026 (sdBV) stars: these are $p$-mode pulsators whose temperatures fall in the range between 29,000 and 36,000~K. Their periods are typically found in the range 100-200~sec, and their amplitudes cover the range from 0.4 to 25\%. \item PG1716+426 (``Betsy'') stars: these are $g$-mode pulsators, with temperatures in the range between 25,000 and 30,000~K. Their periods are much longer, typically falling in the range between 2000 and 9000~sec, and their amplitudes are smaller than 0.5\%. \item Hybrid stars: these are stars that present simultaneous $p$- and $g$- mode oscillations. At present, only two examples have been reported in the literature \citep{abea05,roea05,ssea06}. \end{itemize} Performing asteroseismology on these stars holds the promise to unveil their innermost secrets, thus providing an exciting new route toward the solution of the so-called ``second-parameter problem'' \citep[][and references therein]{mc05}. The great promise of the technique notwithstanding, such variables have never been detected in previous searches in globular clusters \citep[e.g.,][]{mrea06}. Accordingly, the main purpose of the present study is to perform a new search for this type of variables in the relatively nearby southern globular cluster NGC~6752, which contains a very long blue HB ``tail,'' and thus many potential candidates for the three aforementioned variability types. | Our analysis of time-series observations collected with the Magellan 6.5m MagIC camera already reveals a few intriguing variable candidates, including at least one on the hot extension of the HB, in two small fields located away from the cluster center. We have recently acquired extensive time-series observations using IMACS over a much larger field. These data will help confirm the nature of the suspected variables, and will presumably reveal a host of other faint/low-amplitude variables in NGC~6752. In addition, the new data will allow us to map the MS binary fraction in the cluster as a function of radius, which will reveal, for the first time, how in detail the binary fraction decreases as a function of radius in a globular star cluster. | 7 | 10 | 0710.0600 |
0710 | 0710.2433_arXiv.txt | We investigate the wave effect in the gravitational lensing by a black hole with very tiny mass less than $10^{-19} M_{\rm sun}$ (solar mass), which is called attolensing, motivated by a recent report that the lensing signature might be a possible probe of a modified gravity theory in the braneworld scenario. We focus on the finite source size effect and the effect of the relative motion of the source to the lens, which are influential to the wave effect in the attolensing. Astrophysical condition that the lensed interference signature can be a probe of the modified gravity theory is demonstrated. The interference signature in the microlensing system is also discussed. | Many physicists have drawn attention to extra dimensional physics for several years due to recent development in testing Randall-Sundrum type II (RS-II) scenario \cite{RSII}. In this scenario they considered a four dimensional positive-tension brane embedded in five dimensional AdS bulk which allows us to reconsider our understanding about the history of our universe in the early stages. Some investigations have been carried out to modify the existence of primordial black holes (PBH)s in this RS-II scenario that the life time of five dimensional PBHs against the Hawking radiation becomes longer compared with the standard PBH in four dimensions \cite{GCL}. This is because the five dimensional feature becomes significant in the black hole with very tiny mass. The ratio of the life time against the Hawking radiation of such five dimensional PBH to that of the four dimensional PBH may be estimated, $lM_4/l_4M$, where $l$ is the AdS radius of the braneworld model, $l_4$ and $M_4$ are the four dimensional Planck length and mass, respectively, and $M$ is the black hole mass \cite{GCL}. Since the braneworld PBHs can live longer, then it is possible that such black holes still exist and spread out in our universe. Therefore, it is natural to consider the possibility of the gravitational lensing phenomenon by such the black hole. More recently, the authors \cite{KPIII} investigated the wave effect in gravitational lensing by the black hole with the very tiny mass smaller than $10^{-19} M_{\rm sun}$, where $M_{\rm sun}$ is the solar mass, which is called attolensing. They showed that the interference signature in the energy spectrum in gamma ray burst due to the attolensing might be a possible probe of the modified gravity theory. In the standard general relativity, PBH with the mass smaller $10^{-18} M_{\rm sun}$ will be evaporated through the Hawking radiation within the cosmic age. Then, the detection of such the interference signature would be a probe of extra dimension of our universe. In the present paper, we consider the two effects in the lensing phenomenon that are influential for measurement of the interference signature in the energy spectrum in the attolensing: One is the finite source size effect. The other is the effect of the relative motion of the black hole (lens object) to the source. We demonstrate the condition that these effects become influential. This is one of the astrophysical conditions that the attolensing can be a probe of the modified gravity theory. Throughout this paper, $ H_0 = 72$ km s$^{-1}$ Mpc$^{-1}$ is the Hubble parameter, and we use the unit in which the light velocity equals 1. | We investigated the condition that the finite source size effect and the relative motion of the source becomes substantial in the wave effect of the gravitational lensing. The condition is expressed by the formula (\ref{condtwo}), whose physical meaning is that the difference of the phase between the two light paths from $y_{\rm max}$ and $y_{\rm min}$ is larger than $2\pi$. We have shown that the finite source size effect is important in the attolensing by the black hole at the cosmological distance. Also the relative motion of the source to the lens can be influential. For the attolensing by the black hole at the galactic distance, the constraint is relaxed by the factor $\sqrt{D_S/D_L}\sim 10^{3}$. However, we should also note that the detection of the attolensing is limited in practice \cite{KPIII}, even when the finite source size is not taken into account. In general, the signature of the interference becomes remarkable when $w\sim 1$, i.e., \begin{eqnarray} w \sim 0.3\times (1+z_L)\left({h\nu\over 100{\rm MeV}}\right) \left({M\over 10^{-19}M_{\rm sun}}\right)\sim 1. \label{condfirst} \end{eqnarray} In the case $w\simlt 1$, the amplification due to the lensing becomes negligible, because the wavelength of the light becomes larger than the size of deflector. Combining this and the condition (\ref{condtwo}), we may conclude that the condition for the possible observation of the interference signature in the gravitational lensing needs \begin{eqnarray} 1\simlt w \simlt {\pi \over |y_{\rm min}-y_{\rm max}|}. \label{finalcond} \end{eqnarray} Therefore, this means $|y_{\rm min}-y_{\rm max}|\simlt \pi$, the angular size of the source must be less than the Einstein radius, is always necessary for a possible observation of the interference signature. It will be useful to demonstrate the region satisfying (\ref{finalcond}) clearly. In Figure 8, the dashed line is $w=1$, and the solid line is $w|y_{\rm min}-y_{\rm max}| =\pi$, where we fixed $\hat R=10^3$ km and $D_{LS}D_S/D_L=H_0^{-1}$. The shaded region satisfies the condition (\ref{finalcond}). It might also be interesting to consider whether other physical system satisfy the condition (\ref{finalcond}) or not. Figure 9 shows a region satisfying the condition with the parameter associated with the microlensing. Similar to Figure 8, the shaded region satisfies the condition. Here the dashed line in Figure 9 is $w=1$, and the solid line is $w|y_{\rm min}-y_{\rm max}| =\pi$, where we fixed $\hat R=7\times 10^5$ km (solar radius) and $D_{LS}D_S/D_L=30$ kpc. The point of the intersection of these two lines is \begin{eqnarray} && \left({\nu\over {\rm GHz}}\right)=8.9\times 10^2 \left({\hat R\over 7\times 10^5{\rm km}} \right)^{-2} \left({30{\rm kpc}\over D_{LS}D_S/D_L}\right)^{-1} \\ && \left({M\over M_{\rm sun}}\right)=0.9\times 10^{-8} \left({\hat R\over 7\times 10^5{\rm km}} \right)^{2} \left({30{\rm kpc}\over D_{LS}D_S/D_L}\right). \end{eqnarray} Thus the microlensing can be a possible system that satisfies the condition (\ref{finalcond}). Especially, for the microlensing by an Earth like planet $M\sim 10^{-5} M_{\rm sun}$, the relevant range of the frequency is $1$ GHz $\sim 100$ GHz. Then, the measurement of the microlensing event through the frequency might be relevant to the interference signature. However, it will be very difficult to detect the signal because the stars at $1\sim 100$ GHz frequency band at the galactic distance is very dark in general. \vspace{0.3cm} | 7 | 10 | 0710.2433 |
0710 | 0710.5355_arXiv.txt | { The ANTARES Collaboration is building a high-energy neutrino telescope at 2500 m depth in the Mediterranean Sea. The experiment aims to search for high-energy cosmic neutrinos through the detection of Cerenkov light induced by muons and showers resulting from neutrino interactions with the surrounding medium. The detector will consist of a three-dimensional array of 900 optical modules housing photomultipliers. It will be composed of 12 strings, 5 of them being already in operation since January 2007. The muon track is reconstructed from the arrival time and the charge of the signals obtained from the photomultipliers, whose positions are known by means of an acoustic positioning system. The reconstruction strategies include several steps among which there are: optical background filtering, algorithms for first estimations of the track parameters, and a final fit aiming to reach an angular resolution better than 0.3 degree above 10 TeV in the full detector. Different reconstruction strategies will be presented and their application to the present real data analysis will be reviewed.} \begin{document} | When an ultra-relativistic particle ($\beta \simeq 1$) moves in a medium, Cerenkov light is emitted at an angle depending on the refraction index $n$ of the medium. In case of sea water $n$ is about $1.34$, and thus the Cerenkov emission angle becomes \begin{equation} cos (\theta_C) = \frac{1}{n} \simeq cos(42^\circ). \end{equation} The emitted Cerenkov photons are detected by an array of photomultipliers installed at the bottom of the sea. Since January 2007 the ANTARES detector is a full three-dimensional array of photomultipliers consisting of 5 strings detecting muons at a depth of 2475 meters below the sea level. ANTARES 10-inch photomultipliers are housed in pressure resistant glass spheres called optical modules (OM). The 5-line detector data taking has allowed the Collaboration to tune the various processes needed to collect data, and muon events have been detected on top of the optical background (60-100 kHz most of the time). The calibration system has been proven to be efficient, and the knowledge of the charge, of the arrival time of the signals and the positioning of the OMs has allowed the first reconstruction codes to be tested in realistic conditions. The status of the experiment is discussed in \cite{antoine}. Preliminary data studies show that a flux of atmospheric downward-going muons is triggering the detector at a rate of about 1 Hz and that upgoing atmospheric neutrino candidates have been identified. Among the various muon reconstruction algorithms under study, two different alternative methods are presented in this paper. Other well developed reconstruction methods are described in \cite{aart}, \cite{carmona}, and the implementation of the {\sl{simulated annealing}} algorithm \cite{simann} is under study. A discussion on event reconstruction techniques for Cerenkov neutrino telescopes can also be found in \cite{amanda}. | 7 | 10 | 0710.5355 |
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0710 | 0710.5163_arXiv.txt | We report the discovery of 11 new cataclysmic variable (CV) candidates by the Isaac Newton Telescope (INT) Photometric H$\alpha$ Survey of the northern Galactic plane (IPHAS). Three of the systems have been the subject of further follow-up observations. For the CV candidates IPHAS J013031.90+622132.4 and IPHAS J051814.34+294113.2, time-resolved optical spectroscopy has been obtained and radial-velocity measurements of the H$\alpha$ emission-line have been used to estimate their orbital periods. A third CV candidate (IPHAS J062746.41+ 014811.3) was observed photometrically and found to be eclipsing. All three systems have orbital periods above the CV period-gap of 2--3\,h. We also highlight one other system, IPHAS J025827.88+635234.9, whose spectrum distinguishes it as a likely high luminosity object with unusual C and N abundances. | \label{intro} Cataclysmic variables (CVs) are semi-detached interacting binary systems containing a white dwarf (WD) primary and a late-type main-sequence secondary. The secondary fills its Roche lobe, and matter is transferred to the primary through the inner Lagrangian point. The mass-transfer process in CVs means that line emission is a common observational feature of the majority of CVs, especially in the Balmer series. Observed lines may originate from optically thin or irradiated parts of the accretion disc (if present, \citealt{1980ApJ...235..939W}). Line emission may also be produced in the accretion stream and on the irradiated face of the secondary star. \citet{1995cvs..book.....W} provides a comprehensive review of CVs. Population synthesis models suggest that the intrinsically faintest low accretion rate systems should dominate the Galactic CV population and should be found predominantly at short orbital periods, that is, below the well-known CV ``period gap'' between 2 hrs and 3 hrs \citep{1993A&A...271..149K, 1997MNRAS.287..929H}. This dominant population of faint, short-period CVs has proven quite elusive. This is partly because most known CVs have been discovered with techniques that are biased against detecting CVs with low mass accretion rates. For example, all flux-limited surveys with relatively bright limiting magnitudes and also variability searches (e.g. for dwarf nova outbursts) are intrinsically biased against the detection of faint, short-period CVs with potentially long inter-outburst recurrence times (such as WZ Sge). For a closer look at the period distributions of CVs found with different discovery methods, see \citet{2005ASPC..330....3G}. Recently \citet{2007MNRAS.374.1495P} have shown that the relative dearth of short-period CVs is not just due to selection effects, implying a serious flaw in our understanding of CV evolution (see also \citealt{2006A&A...455..659A}). Nevertheless, selection effects must be at least partly responsible for the lack of known faint short-period CVs, and so many such systems still remain to be found. Given the ubiquity of line emission amongst CVs, searches for objects displaying H$\alpha$ emission offer a powerful way to find new CVs. Examples of such CV searches in the Galactic field include those of \citet{2002ASPC..261..190G}, \citet{2005A&A...443..995A}, \citet{2002A&A...395L..47H}, and the Chandra Multiwavelength Plane (ChaMPlane) survey (\citealt{2005ApJ...635..920G}, \citealt{2005ApJS..161..429Z} and \citealt{2006ApJS..163..160R}). Similar searches have also been performed in globular clusters have been performed by \citet{1995ApJ...439..695C}, \citet{1995ApJ...455L..47G}, \citet{1996ApJ...473L..31B} and \citet{2000ApJ...532..461C}. The Isaac Newton Telescope (INT) Photometric H$\alpha$ Survey of the northern Galactic plane (IPHAS) is currently surveying the Milky Way in broad-band Sloan $r^\prime$ and $i^\prime$ and narrow-band H$\alpha$ and provides an excellent data base for a detailed CV search at low Galactic latitudes. The survey goes to a depth of $r^\prime\simeq20$\,mag and covers the latitude range $-5^o < b < +5^o$. A detailed introduction to the survey, including the transmission profiles of the filters it uses, is given by \citet{2005MNRAS.362..753D}. What makes IPHAS particularly promising for finding CVs is that, empirically, the intrinsically faintest, low mass transfer rate ($\dot{M}$) systems tend to have the largest Balmer line equivalent widths (EWs; \citealt{1984ApJS...54..443P}; \citealt{2006MNRAS.369..581W}). Unless the inverse relationship between Balmer EW and orbital period breaks down for the faintest, short-period systems, emission-line surveys should be very good at finding low-$\dot{M}$ CVs. IPHAS should then be an excellent way of detecting a large population of faint, short-period CVs. In this paper we announce the discovery of eleven new CV candidates from the IPHAS survey data. We present spectroscopic follow-up observations of two of these CVs candidates and determine their orbital periods. We also present the results of time-series photometry of another CV candidate which establishes it as a long period, eclipsing, system. | \label{conc} Spectroscopy of several objects from the IPHAS H$\alpha$ excess catalogue has led to the discovery of 11 new CV candidates by virtue of their strong H$\alpha$ emission. The identification spectra of the CVs candidates include cases where the secondary star is particularly prominent (IPHAS J1856), and examples of double-peaked emission lines (for example IPHAS J0518). One particular candidate (IPHAS J0258) has similarities to the extreme objects V Sge and QU Car. Time-resolved optical spectroscopy of two of the new CVs (IPHAS J0130 and IPHAS J0518) has been obtained using CAFOS on the 2.2\,m telescope at the Calar Alto Observatory. Radial-velocity measurements of the H$\alpha$ emission-line were used to determine their orbital periods. Periodograms and a Monte Carlo analysis were used to estimate the orbital period of $P_{orb}=0.130061 \pm 0.000001$\,d for IPHAS J0130. The periodogram obtained from the radial-velocity data of IPHAS J0518 reveals four peaks which could be the orbital period of the system: 0.2383, 0.2203, 0.2595 and 0.2049\,d, each with an uncertainty of 0.0007\,d. Time-series photometry of IPHAS J0627 using the SAAO 1.9\,m telescope found the system to be a long-period eclipser with four possible orbital periods: $1.020 \pm 0.002$\,d, $0.5101 \pm 0.0008$\,d, $0.3401 \pm 0.0006$\,d and $0.2551 \pm 0.0004$\,d. Further observations are necessary to obtain accurate values of the orbital period of all three systems, and to find accurate binary parameters. | 7 | 10 | 0710.5163 |
0710 | 0710.4954_arXiv.txt | We study the signal for the detection of quasi-stable supersymmetric particle produced in interactions of cosmogenic neutrinos. We consider energy loss of high energy staus due to photonuclear and weak interactions. We show that there are optimal nadir angles for which the stau signal is a factor of several hundred larger than muons. We discuss how one could potentially eliminate muon background by considering the energy loss of muons in the detector. We also show results for the showers produced by weak interactions of staus that reach the detector. | Ultrahigh energy cosmic neutrinos could potentially probe physics beyond the Standard Model \cite{ringwald}. Interactions of UHE neutrinos ($E_\nu \geq 10^{17}$eV) with nucleons probe center of mass energies above 14 TeV. Some fraction of these neutrinos may produce supersymmetric particles or some other exotic particles. These processes are suppressed relative to standard model processes, however, in some models interesting signals may arise from supersymmetric particles with long lifetimes. In most SUSY scenarios, particles produced in high energy collisions decay immediately into the lightest one and are thus hard to detect. However in some low scale supersymmetric models in which gravitino is the lightest supersymmetric particle (LSP) and R-parity is conserved, the next-to-lightest particle (NLSP) is the charged superpartner of the right-handed tau, the stau \cite{susy_models}. Due to its weak coupling to the gravitino, the stau is a long-lived particle in these models. For the supersymmetric breaking scale, $\sqrt F >5\times10^6$ GeV, the long-lived stau could travel distances of the order of $10^4$ km before decaying into the gravitino. The distance that staus travel before decaying depends on the gravitino mass (or equivalently on the supersymmetry breaking scale) and the stau mass. Limits on the stau mass of about 100 GeV come from its non-observation in accelerator experiments \cite{stmass1,stmass2,stmass3,stmass4}. Recently it was proposed that staus produced in high energy neutrino interactions, where neutrinos originate in astrophysical sources, might be detectable in neutrino telescopes \cite{ABC,Ahlers}. The cross section for the production of staus in neutrino-nucleon scattering \cite{ABC} is several orders of magnitude smaller than the neutrino charged-current or neutral-current cross section \cite{gqrs}. However, once produced, the long-lived staus have the potential to travel through the earth without decaying and thus open up a possibility to be detected in neutrino telescopes. The long range of staus could potentially compensate for the suppression in the production cross section by increasing the effective detector volume and therefore enhancing the signal. The detection of staus depends on the stau lifetime and range, so it is important to determine the energy loss as it traverses the earth. The details of the range depend in part on the supersymmetry breaking and how the quasi-stable stau particle is comprised of the SUSY partners of the right-handed and left-handed taus. The electromagnetic energy loss has been shown to have the largest contribution from photonuclear interactions for stau energies between $10^6$-$10^{12}$ GeV, resulting in a range of $10^4$ km.w.e. for masses of the order of a few hundred GeV \cite{RSS}. Weak interactions may come into play as well. The stau range has been shown to be sensitive to the mixing angle of right-handed and left-handed staus. When the mixing is maximal, weak interactions act to suppress the range at energies above $\sim10^9$ GeV \cite{RSU1}, however, their weak interactions have the potential to produce signals in neutrino detectors such as the Antarctic Impulse Transient Array (ANITA) \cite{anita} and the Antarctic Ross Iceshelf Antenna Neutrino Array (ARIANNA) \cite{arianna}. The high energies required for stau production lead us to focus on the production of staus in interactions of cosmogenic neutrinos as they traverse the Earth and/or in the detector. These neutrinos originate from cosmic ray protons interacting with the cosmic microwave background, $$ p\gamma(3{\rm K}) \rightarrow \Delta \rightarrow N\pi$$ followed by charged pion, muon and neutron decays. This flux is guaranteed as cosmic ray fluxes are measured as well as the 3K microwave background. We use a conservative cosmogenic neutrino flux evaluated by Engel, Seckel and Stanev (ESS) in Ref. \cite{ess}. They evaluate the neutrino flux associated with the measured cosmic ray flux by tracing back cosmic ray propagation through the background radiation. Depending on the cosmological evolution assumed, the overall normalization of this flux has an uncertainty of about a factor of four. In addition, neutrinos could be produced at the sources of the high energy cosmic rays and those are not included in the evaluation of ESS neutrino flux. Thus, the ESS neutrino flux is a conservative estimate of the cosmogenic flux. The cosmogenic neutrino flux, when neutrino flavor oscillations are not included, peaks at high energies, around $10^8$ GeV, and thus it is in the energy range where ANITA and ARIANNA have very good sensitivity \cite{anita,arianna}. The neutrino flavor ratio for cosmogenic neutrinos deviates from the common 2:1 ratio, due to the neutron decay contribution to electron neutrinos \cite{ess}. This implies that one needs to consider three flavor oscillations when considering the cosomogenic neutrino flux that arrives at the Earth \cite{us_JMRS}. We consider cosmogenic neutrinos, their propagation, stau production, and subsequent energy loss as it traverses the earth, for a region of parameter space where the staus do not decay over the distances required. We compare the resulting stau flux when there is no mixing between right-handed and left-handed stau and when there is maximal mixing. We consider muon-like signals (charged tracks) produced by staus and its associated background. We discuss the potential for eliminating the background by measuring the energy loss, which requires large volume detectors. Finally we discuss the showers produced in the ice due to stau interactions and its background from neutrino-induced showers. | We have studied signals of staus produced in interactions of cosmogenic neutrinos. We have considered two types of signals, muon-like charged tracks and showers. We have focused on low scale supersymmetric models that have stau as NLSP, which decays into the lightest SUSY particle, the gravitino. For a sufficiently large scale of supersymmetry breaking, the stau has a very long lifetime. Our focus has been on cosmogenic neutrino fluxes and their associated stau production in the Earth. Energy losses, both through electromagnetic and weak interactions, are important in evaluating stau signals. The energy loss of staus, however, is relatively small in comparison with muons. Thus, for some nadir angles, the stau flux is much larger than the muon flux produced in neutrino charged-current interactions. The enhancement of the stau flux is larger from an input cosmogenic neutrino flux than for the Waxman-Bahcall neutrino flux \cite{WB}. This is because the cosmogenic neutrino flux is peaked at energies of about $10^8$ GeV, while the WB flux is characterized by a steep power law with index of two. The large ratio of staus to muons from cosmogenic neutrinos is encouraging for experimental detection, but in order to see this signal one needs to be able to distinguish between staus and muons. Using the average energy loss per unit distance is not a good way to distinguish staus and muons, since the scaling of the energy loss parameter $\beta$ has the effect of making a high-energy stau look like a lower energy muon. We have proposed a way to distinguish between stau and muon tracks by measuring the energy loss of muons via their interactions in the ice, and to use this method to reduce the background. We also considered showers produced by staus interacting in the ice via charged-current interactions. The backgrounds for this signal are showers induced directly by neutrinos that reach the detector and interact inside the detector via charged-current or neutral-current interactions. The only way that staus would produce showers in the ice is if there is a weak mixing, but this process also contributes to reducing stau range. These effects combine with the small stau production probability to give fluxes of attenuated staus that are several orders of magnitude less than attenuated neutrino fluxes. This small stau to neutrino ratio translates directly to shower rates. In addition to weak interactions, another possibility for shower production would be stau decays in the detector. For the parameter space considered here, with long-lived staus, the signal from decays is suppressed relative to their weak interactions. In conclusion, stau signals at high energies are best identified by muon-like tracks. The most important detection issue is distinguishing between staus and muons, which may be possible by looking at the incremental electromagnetic energy loss as the charged particle moves through the detection volume. With very large volumes, there is a potential for detection of staus with future large neutrino telescopes. | 7 | 10 | 0710.4954 |
0710 | 0710.3868_arXiv.txt | Two dimensional hydrodynamical disks are nonlinearly unstable to the formation of vortices. Once formed, these vortices essentially survive forever. What happens in three dimensions? We show with incompressible shearing box simulations that in 3D a vortex in a short box forms and survives just as in 2D. But a vortex in a tall box is unstable and is destroyed. In our simulation, the unstable vortex decays into a transient turbulent-like state that transports angular momentum outward at a nearly constant rate for hundreds of orbital times. The 3D instability that destroys vortices is a generalization of the 2D instability that forms them. We derive the conditions for these nonlinear instabilities to act by calculating the coupling between linear modes, and thereby derive the criterion for a vortex to survive in 3D as it does in 2D: {\it the azimuthal extent of the vortex must be larger than the scale height of the accretion disk}. When this criterion is violated, the vortex is unstable and decays. Because vortices are longer in azimuthal than in radial extent by a factor that is inversely proportional to their excess vorticity, a vortex with given radial extent will only survive in a 3D disk if it is sufficiently weak. This counterintuitive result explains why previous 3D simulations always yielded decaying vortices: their vortices were too strong. Weak vortices behave two-dimensionally even if their width is much less than their height because they are stabilized by rotation, and behave as Taylor-Proudman columns. We conclude that in protoplanetary disks weak vortices can trap dust and serve as the nurseries of planet formation. Decaying strong vortices might be responsible for the outwards transport of angular momentum that is required to make accretion disks accrete. | Matter accretes onto a wide variety of objects, such as young stars, black holes, and white dwarfs, through accretion disks. In highly ionized disks magnetic fields are important, and they trigger turbulence via the magnetorotational instability \citep{BH98}. However, many disks, such as those around young stars or dwarf novae, are nearly neutral \citep[e.g.,][]{SMUN00,GM98}. In these disks, the fluid motions are well described by hydrodynamics. Numerical simulations of hydrodynamical disks in two-dimensions---in the plane of the disk---often produce long-lived vortices \citep{GL99,UR04,JG05}. If vortices really exist in accretion disks, they can have important consequences. First and foremost, they might generate turbulence. Since turbulence naturally transports angular momentum outwards\footnote{ Energy conservation implies that turbulence transports angular momentum outwards; see \S \ref{sec:pseudo}. Nonetheless, if an external energy source (e.g., the radiative energy from the central star) drives the turbulence, then angular momentum could in principle be transported inwards. }, as is required for mass to fall inwards, it might be vortices that cause accretion disks to accrete. Second, in disks around young stars, long-lived vortices can trap solid particles and initiate the formation of planets \citep{BS95}. Why do vortices naturally form in 2D simulations? Hydrodynamical disks are stable to linear perturbations. However, they are nonlinearly unstable, despite some claims to the contrary in the astrophysical literature. In two dimensions, the incompressible hydrodynamical equations of a disk are equivalent to those of a non-rotating linear shear flow \citep[e.g.,][hereafter L07]{L07}. And it has long been known that such flows are nonlinearly unstable (\mycite{Gill65}; \mycite{LK88}; L07). This nonlinear instability is just a special case of the Kelvin-Helmholtz instability. Consider a linear shear flow extending throughout the $x$-$y$ plane with velocity profile $\bld{v}=-qx\bld{\hat{y}}$, where $q>0$ is the constant shear rate, so that $-q$ is the flow's vorticity. (In the equivalent accretion disk, the local angular speed is $\Omega=2q/3$.) This shear flow is linearly stable to infinitesimal perturbations. But if the shear profile is altered by a small amount, the alteration can itself be unstable to infinitesimal perturbations. To be specific, let the alteration be confined within a band of width $\Delta x$, and let it have vorticity $\omega=\omega(x)$ (with $|\omega|\lesssim q$), so that it induces a velocity field in excess of the linear shear with components $u_y\sim \omega\Delta x$ and $u_x=0$. Then this band is unstable to infinitesimal nonaxisymmetric (i.e. non-stream-aligned) perturbations provided roughly that \be \left|k_y \right| \lesssim {1\over q}{|\omega|\over \Delta x} \ \ \Rightarrow {\rm 2D\ instability} \label{eq:2dinst} \ee where $k_y$ is the wavenumber of the nonaxisymmetric perturbation.\footnote{ More precisely, the necessary and sufficient condition for instability in the limit $|\omega|\ll q$ is that $|k_y|<{1\over 2 q}\int_{-\infty}^\infty {d\omega/dx\over x-x_0}dx$, where $x_0$ is any value of $x$ at which $d\omega/dx=0$ \citep[][L07]{Gill65,LK88}. For arbitrarily large $\omega$, Rayleigh's inflection point theorem and Fj\o rtoft's theorem give necessary (though insufficient) criteria for instability \citep{DR04}. The former states that for instability, it is required that $d\omega/dx=0$ somewhere in the flow, i.e. that the velocity field must have an inflection point. \cite{Love99} generalize Rayleigh's inflection point theorem to compressible and nonhomentropic disks. \label{foot:inst} } For any value of $|\omega|$ and $\Delta x$, the band is always unstable to perturbations with long enough wavelength. Remarkably, instability even occurs when $|\omega|$ is infinitesimal. Hence we may regard this as a true nonlinear instability. \cite{BH06} assert that detailed numerical simulations have not shown evidence for nonlinear instability. The reason many simulations fail to see it is that their boxes are not long enough in the $y$-direction to encompass a small enough non-zero $|k_y|$. In two dimensions, the outcome of this instability is a long-lived vortex (e.g., L07). A vortex that has been studied in detail is the Moore-Saffman vortex, which is a localized patch of spatially constant vorticity superimposed on a linear shear flow \citep{Saffman95}. When $|\omega|\lesssim q$, where $\omega$ here refers to the spatially constant excess vorticity within the patch, and when the vorticity within the patch ($\omega-q$) is stronger than that of the background shear, then the patch forms a stable vortex that is elongated in $y$ relative to $x$ by the factor \be {\Delta y\over \Delta x} \sim {q\over |\omega|} \ . \label{eq:ms} \ee This relation applies not only to Moore-Saffman vortices, but also to vortices whose $\omega$ is not spatially constant. It may be understood as follows. A patch with characteristic excess vorticity $\sim\omega$ and with $\Delta y\gg \Delta x$ induces a velocity field in the $x$-direction with amplitude $u_x\sim|\omega|\Delta x$, independent of the value of $\Delta y$ (e.g., \S 6 in L07). As long as $|\omega|\lesssim q$, the $y$-velocity within the vortex is predominantly due to the background shear, and is $\sim q\Delta x$. Therefore the time to cross the width of the vortex is $t_x\sim \Delta x/u_x \sim 1/|\omega|$, and the time to cross its length is $t_y\sim \Delta y/(q\Delta x)$. Since these times must be comparable in a vortex, equation (\ref{eq:ms}) follows. Equation (\ref{eq:ms}) is very similar to equation (\ref{eq:2dinst}). The 2D instability naturally forms into a 2D vortex. Futhermore, the exponential growth rate of the instability is $\sim |\omega|$, which is comparable to the rate at which fluid circulates around the vortex. More generally, an arbitrary axisymmetric profile of $\omega(x)$ tends to evolve into a distinctive banded structure. Roughly speaking, bands where $\omega<0$ contain vortices, and these are interspersed with bands where $\omega>0$, which contain no vortices. (Recall that we take the background vorticity to be negative; otherwise, the converse would hold.) The reason for this is that only regions that have $\omega<0$ can be unstable, as may be inferred either from the integral criterion for instability given in footnote \ref{foot:inst}, or from Fj\o rtoft's theorem. For more detail on vortex dynamics in shear flows, see the review by \cite{Marcus93}. What happens in three dimensions? To date, numerical simulations of vortices in 3D disks have been reported in two papers. \cite{BM05b} initialized their simulation with a Moore-Saffman vortex, and solved the anelastic equations in a stratified disk. They found that this vortex decayed. As it decayed, new vortices were formed in the disk's atmosphere, two scale heights above the midplane. The new vortices survived for the duration of the simulation. \cite{SSG06} performed both 2D and 3D simulations of the compressible hydrodynamical equations in an unstratified disk, initialized with large random fluctuations. They found that whereas the 2D simulations produced long-lived vortices, in three dimensions vortices rapidly decayed. Intuitively, it seems clear that a vortex in a very thin disk will behave as it does in 2D. And from the 3D simulations described above it may be inferred that placing this vortex in a very thick disk will induce its decay. Our main goal in this paper is to understand these two behaviors, and the transition between them. A crude explanation of our final result is that vortices decay when the 2D vortex motion couples resonantly to 3D modes, i.e., to modes that have vertical wavenumber $k_z\ne 0$. As described above, a vortex with excess vorticity $|\omega|$ has circulation frequency $\sim|\omega|$, and $k_y/k_x\sim |\omega|/q$, where $k_x$ and $k_y$ are its ``typical'' wavenumbers. Furthermore, it is well-known that the frequency of axisymmetric ($k_y=0$) inertial waves is $\Omega k_z/\sqrt{k_x^2+k_z^2}$ (see eq. [\ref{eq:axi3d}]). Equating the two frequencies, and taking the $k_x$ of the 3D mode to be comparable to the $k_x$ of the vortex, as well as setting $q=3\Omega/2$ for a Keplerian disk, we find \be k_z\sim k_y \label{eq:res} \ee as the condition for resonance. Therefore a vortex with length $\Delta y$ will survive in a box with height $\Delta z\lesssim \Delta y$, because in such a box all 3D modes have too high a frequency to couple with the vortex, i.e., all nonzero $k_z$ exceed the characteristic $k_y\sim 1/\Delta y$. But when $\Delta z\gtrsim \Delta y$, there exist $k_z$ in the box that satisfy the resonance condition (\ref{eq:res}), leading to the vortex's destruction. This conclusion suggests that vortices live indefinitely in disks with scale height less than their length ($h\lesssim \Delta y$) because in such disks all 3D modes have too high a frequency for resonant coupling. This conclusion is also consistent with the simulations of \cite{BM05b} and \cite{SSG06}. Both of these works initialized their simulations with strong excess vorticity $|\omega|\sim q$, corresponding to nearly circular vortices. Both had vertical domains that were comparable to the vortices' width. Therefore both saw that their vortices decayed. Had they initialized their simulations with smaller $|\omega|$, and increased the box length $L_y$ to encompass the resulting elongated vortices, both would have found long-lived 3D vortices. \citeauthor{BM05b}'s discovery of long-lived vortices in the disk's atmosphere is simple to understand because the local scale height is reduced in inverse proportion to the height above the midplane. Therefore higher up in the atmosphere the dynamics becomes more two-dimensional, and a given vortex is better able to survive the higher it is.\footnote{ However, \cite{BM05b} also include buoyancy forces in their simulations, which we ignore here. How buoyancy affects the stability of vortices is a topic for future work. } \subsection{Organization of the Paper} In \S \ref{sec:eom} we introduce the equations of motion, and in \S \ref{sec:pseudo} we present two pseudospectral simulations. One illustrates the formation and survival of a vortex in a short box, and the other illustrates the destruction of a vortex in a tall box. In \S\S \ref{sec:lin}-\ref{sec:nonlin} we develop a theory explaining this behavior. The reader who is satisfied by the qualitative description leading to equation (\ref{eq:res}) may skip those two sections. The theory that we develop is indirectly related to the transient amplification scenario for the generation of turbulence. Even though hydrodynamical disks are linearly stable, linear perturbations can be transiently amplified before they decay, often by a large factor. It has been proposed that sufficiently amplified modes might couple nonlinearly, leading to turbulence \citep[e.g.,][]{CZTL03,Yecko04,AMN05}. However, to make this proposal more concrete, one must work out how modes couple nonlinearly. In L07, we did that in two dimensions. We showed that the 2D nonlinear instability of equation (\ref{eq:2dinst}) is a consequence of the coupling of an axisymmetric mode with a transiently amplified mode, which may be called a ``swinging mode'' because its phasefronts are swung around by the background shear. In \S \ref{sec:nonlin} we show that the 3D instability responsible for the destruction of vortices is a generalization of this 2D instability. It may be understood by examining the coupling of a 3D swinging mode with an axisymmetric one. 3D modes become increasingly unstable as $|k_z|$ decreases, and in the limit that $k_z\rightarrow 0$, the 3D instability matches smoothly onto the 2D one. Thicker disks are more prone to 3D instability because they encompass smaller $|k_z|$. | \label{sec:conc} Our main result follows from Figure \ref{fig:marg}, which maps out the stability of a ``mother mode'' (i.e., a mode with wavevector $\bar{k}\bld{\hat{x}}$ and amplitude $\bar{\omega}$) to nonaxisymmetric 3D perturbations. A mother mode is unstable provided that the $k_y$ and $k_z$ of the nonaxisymmetric perturbations satisfy both $|k_y|\lesssim \bar{k}\bar{\omega}/q$ and $|k_z|\lesssim |k_y|$, dropping order-unity constants. Based on this result, we may understand the formation, survival, and destruction of vortices. Vortices form out of mother modes that are unstable to 2D ($k_z=0$) perturbations. Mother modes that are unstable to 2D modes but stable to 3D ($k_z\ne 0$) ones, form into long-lived vortices. Mother modes that are unstable to both 2D and 3D modes are destroyed. Therefore a mother mode with given $\bar{k}$ and $\bar{\omega}$ will form into a vortex if the disk has a sufficiently large circumferential extent and a sufficiently small scale height, i.e., if $r\gtrsim \bar{k}^{-1}q/\bar{\omega}$ and $h\lesssim \bar{k}^{-1}q/\bar{\omega}$, where $r$ is the distance to the center of the disk, and $h$ is the scale height. Alternatively, the mother mode will be destroyed in a turbulent-like state if both $r$ and $h$ are sufficiently large ($r\gtrsim \bar{k}^{-1}q/\bar{\omega}$ and $h\gtrsim \bar{k}^{-1}q/\bar{\omega}$). Our result has a number of astrophysical consequences. In protoplanetary disks that do not contain any vortices, solid particles drift inward. Gas disks orbit at sub-Keplerian speeds, $v_{\rm gas}\sim \Omega r (1-\eta)$, where $\Omega r$ is the Keplerian speed and $\eta\sim (c_s/\Omega r)^2$, with $c_s$ the sound speed. Since solid particles would orbit at the Keplerian speed in the absence of gas, the mismatch of speeds between solids and gas produces a drag on the solid particles, removing their angular momentum and causing them to fall into the star. For example, in the minimum mass solar nebula, meter-sized particles fall in from 1 AU in around a hundred years. This rapid infall presents a serious problem for theories of planet formation, since it is difficult to produce planets out of dust in under a hundred years. Vortices can solve this problem \citep{BS95}. A vortex that has excess vorticity $-\bar{\omega}$ and radial width $1/\bar{k}$ can halt the infall of particles provided that $\bar{\omega}/\bar{k}\gtrsim (\Omega r)\eta$, because the gas speed induced by such a vortex more than compensates for the sub-Keplerian speed induced by gas pressure.\footnote{ We implicitly assume here that the stopping time of the particle due to gas drag is comparable to the orbital time, which is true for meter-sized particles at 1 AU in the minimum mass solar nebula. A more careful treatment shows that a vortex can stop a particle with stopping time $t_s$ provided that $\bar{\omega}/\bar{k}\gtrsim (\Omega r)(\Omega t_s)\eta$ \citep{Youdin08}. } Previous simulations implied that 3D vortices rapidly decay, and so cannot prevent the rapid infall of solid particles \citep{BM05b,SSG06}. Our result shows that vortices can survive within disks, and so restores the viability of vortices as a solution to the infall problem. A more important---and more speculative---application of our result is to the transport of angular momentum within neutral accretion disks. In our simulation of a vortex in a tall box, we found that as the vortex decayed it transported angular momentum outward at a nearly constant rate for hundreds of orbital times. If decaying vortices transport a significant amount of angular momentum in disks, they would resolve one of the most important outstanding questions in astrophysics today: what causes hydrodynamical accretion disks to accrete? To make this speculation more concrete, one must understand the amplitude and duration of the ``turbulence'' that results from decaying vortices. This is a topic for future research. In this paper, we considered only the effects of rotation and shear on the stability of vortices, while we neglected the effect of vertical gravity. There has been a lot of research in the geophysical community on the dynamics of fluids in the presence of vertical gravity, since stably stratified fluids are very common on Earth---in the atmosphere, oceans, and lakes. In numerical and laboratory experiments of strongly stratified flows, thin horizontal ``pancake vortices'' often form, and fully developed turbulence is characterized by thin horizontal layers. \citep[e.g., ][]{BBLC07}. Pancake vortices are stabilized by vertical gravity, in contrast to the vortices studied in this paper which are stabilized by rotation. Gravity inhibits vertical motions because of buoyancy: it costs gravitational energy for fluid to move vertically. The resulting quasi-two-dimensional flow can form into a vortex.\footnote{\cite{BC00} show that a vertically uniform vortex column in a stratified (and non-rotating and non-shearing) fluid suffers an instability (the ``zigzag instability'') that is characterized by a typical vertical lengthscale $\lambda_z\sim U/N$, where $U$ is the horizontal speed induced by the vortex, $N$ is the Brunt-V\"ais\"al\"a frequency, and the horizontal lengthscale of the vortex $L_h$ is assumed to be much greater that $\lambda_z$ (hence the pancake structure). We may understand Billant \& Chomaz's result in a crude fashion with an argument similar to that employed in the introduction to explain the destruction of rotation-stabilized vortices: since the frequency of buoyancy waves is $Nk_x/k_z$ (when $|k_x|\ll |k_z|$), and since the frequency at which fluid circulates around a vortex is $U/L_h\sim k_xU$, there is a resonance between these two frequencies for vertical lengthscale $1/k_z\sim U/N$.} We may speculate that in an astrophysical disk vertical gravity provides an additional means to stabilize vortices, in addition to rotation. But to make this speculation concrete, the theory presented in this paper should be extended to include vertical gravity. We have not addressed in this paper the origin of the axisymmetric structure (the mother modes) that give rise to surviving or decaying vortices. One possibility is that decaying vortices can produce more axisymmetric structure, and therefore they can lead to self-sustaining turbulence. This seems to us unlikely. We have not seen evidence for it in our simulations, but this could be because of the modest resolution of our simulations. Other possibilities for the generation of axisymmetric structure include thermal instabilities, such as the baroclinic instability, or convection, or stirring by planets. This, too, is a topic for future research. \appendix | 7 | 10 | 0710.3868 |
0710 | 0710.3059_arXiv.txt | \label{sec:discussion} We considered the relation between the HMXB population and the SFH of the host galaxy. The number of HMXBs can be represented as a convolution (Eq. (4)) of the star formation history SFR(t) with the function $\eta_{HMXB}(t)$ describing the dependence of the HMXB number on the time elapsed since the star formation event. Thus, the evolution of the HMXB population after the star formation event can be reconstructed by analyzing the distribution of HMXBs in stellar complexes with different SFHs. Using archival optical observations, we reconstructed the spatially resolved SFH in the SMC over the past $\sim$100~Myr (Fig. 10). For this purpose, the observed color-magnitude diagrams of the stellar population were approximated by linear combinations of model isochrones. We analyzed the stability and errors of this method for reconstructing the recent SFH and showed that its accuracy is limited by the uncertainties in the currently available models for the evolution of massive stars. However, the systematic error introduced by this factor may be ignored, since the main source of uncertainty in the solution is the Poisson noise due to the relatively small number of HMXBs in the part of the SMC investigated by XMM-Newton. Using the derived SFHs and the spatial distribution of HMXBs in the SMC from Shtykovskiy and Gilfanov (2005b), we reconstructed the function $\eta_{HMXB}(t)$ that describes the dependence of the HMXB number on the time elapsed since the star formation event (Fig. 12). We compared the derived dependence with the behavior of the SN II rate. The HMXB number reaches its maximum $\sim$20--50~Myr after the star formation event, which is comparable to or exceeds the lifetime of a $8M_\odot$ star. This is much later than the maximum of the SN II rate. In addition, note the shortage of the youngest systems. Observationally, this manifests itself in the absence (or an extremely small number) of HMXBs with black holes in the SMC. This behavior is related to the evolution of the companion star and the neutron star spin period and is consistent with the population synthesis model calculations (Popov et al. 1998). When these results are interpreted, it should be kept in mind that the function $\eta_{HMXB}(t)$ depends on the luminosity threshold used to select the X-ray sources. In our analysis, we used a sample with a low luminosity threshold, L$_{min}\sim 10^{34}$~erg/s. In such a sample, low-luminosity sources, mostly Be/X systems, mainly contribute to the number of sources, while the relative contribution from systems with black holes and/or O/B supergiants, which must constitute the majority of sources in the lower time bin in Fig. 12, is small. Therefore, the time dependence of the number of bright sources (e.g., $>10^{37}$~erg/s) will differ from that shown in Fig. 12. The HMXB formation efficiency in the SMC does not exceed the prediction of the N$_{HMXB}$--SFR calibration (Grimm et al. 2003). Their abnormal abundance compared to the predictions based on the emission in standard SFR indicators, such as the H$_{\alpha}$ line, can result from a peculiarity of the SFH in the SMC. | 7 | 10 | 0710.3059 |
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0710 | 0710.4509_arXiv.txt | We present a scheme to solve the nonlinear multigroup radiation diffusion (MGD) equations. The method is incorporated into a massively parallel, multidimensional, Eulerian radiation-hydrodynamic code with adaptive mesh refinement (AMR). The patch-based AMR algorithm refines in both space and time creating a hierarchy of levels, coarsest to finest. The physics modules are time-advanced using operator splitting. On each level, separate ``level-solve'' packages advance the modules. Our multigroup level-solve adapts an implicit procedure which leads to a two-step iterative scheme that alternates between elliptic solves for each group with intra-cell group coupling. For robustness, we introduce pseudo transient continuation ($\ptc$). We analyze the magnitude of the $\ptc$ parameter to ensure positivity of the resulting linear system, diagonal dominance and convergence of the two-step scheme. For AMR, a level defines a subdomain for refinement. For diffusive processes such as MGD, the refined level uses Dirichet boundary data at the coarse-fine interface and the data is derived from the coarse level solution. After advancing on the fine level, an additional procedure, the sync-solve (SS), is required in order to enforce conservation. The MGD SS reduces to an elliptic solve on a combined grid for a system of $G$ equations, where $G$ is the number of groups. We adapt the ``partial temperature'' scheme for the SS; hence, we reuse the infrastructure developed for scalar equations. Results are presented. We consider a multigroup test problem with a known analytic solution. We demonstrate utility of $\ptc$ by running with increasingly larger timesteps. Lastly, we simulate the sudden release of energy $Y$ inside an Al sphere ($r = 15$ cm) suspended in air at STP\@. For $Y = 11$ kT, we find that gray radiation diffusion and MGD produce similar results. However, if $Y = 1$ MT, the two packages yield different results. Our large $Y$ simulation contradicts a long-standing theory and demonstrates the inadequacy of gray diffusion. | \label{intro} This paper describes a numerical method to solve the radiation multigroup diffusion (MGD) equations. Two themes are presented. One is the scheme itself. We add Pseudo Transient Continuation $(\ptc)$ to the familiar ``fully implicit'' method of Axelrod et al \cite{AxDuRh}. The second theme is code-specific. Our MGD solver is embedded in a multidimensional, massively parallel, Eulerian radiation-hydrodynamic code, which has patch-based, time-and-space Adaptive Mesh Refinement (AMR) capability. Our code's AMR framework stems from the Berger and Oliger idea \cite{BeOl} developed for hyperbolic, compressible hydrodynamic schemes. The idea was expanded by Almgren et al \cite{Alm} and applied to the type of elliptic solvers required for the incompressible equations of Navier-Stokes. Howell and Greenough \cite{HowGre} applied the Almgren et al framework to the scalar, parabolic ``gray'' radiation diffusion equation, thereby creating the start of our radiation-hydrodynamic code. The AMR framework works as follows. A domain, referred to as the ``coarse'' or L0 level, is discretized using a uniform, coarse spatial mesh size $h_c$.\footnote{In multiple dimensions, coordinates have their own mesh spacing.} After advancing with a timestep $\dtc$, the result is scanned for possible improvement. One may refine subregions containing a chosen material, at material interface(s), or at shocks, etc. Whatever refinement criteria are used, after the subdomains are identified, specific routines define a collection of ``patches,'' which cover the subdomains. In two dimensions, the patches are unions of rectangles; in 3D, they are unions of hexahedra. The patches need not be connected, but they must be contained within the coarse level. The patches denote the ``fine'' or L1 level and are discretized with a uniform, spatial mesh size $h_f$. A typical refinement ratio $h_c/h_f$ equals two, but higher multiples of two are also allowed. Because the original framework was designed for temporally explicit hyperbolic schemes, $\dtc$ is restricted by a CFL condition. This implies a similar restriction for the L1 level timestep $\dtf$. For the case, $h_c/h_f = 2$, level L1 time-advances twice using $\dtf = \dtc / 2$. Boundary conditions for level L1 are supplied as follows. Wherever level L1 extends to the physical boundary, the level uses the conditions prescribed by the problem. Portions of level L1's boundary which lie inside the physical domain have conditions prescribed by time and space interpolated data obtained from the L0 solution. For diffusion equations, these conditions are of Dirichlet type. The numerical solution consists of both coarse and fine grid results. Unfortunately, as it stands, the composite solution does not guarantee conservative fluxes across the level boundaries. To maintain conservation, a separate procedure, dubbed a sync-solve (SS) is required. The SS reduces to an elliptic unstructured grid solve on the composite grid of L0 and L1 levels. The AMR procedure may be recursive. That is, a level L1 grid may generate its own subdomain for refinement, i.e., a level L2. In that case, one SS couples results from levels L1 and L2. Once the levels advance to the L0 level time, a SS coupling all three levels ensues. For the multigroup equations, the SS requires an unstructured grid solve for a coupled system of reaction-diffusion equations. Our scheme for a multigroup SS is an important theme of this paper. The MGD equations stem from a discretization of the multifrequency radiation diffusion equations. The latter is an approximation to the equations of radiation transfer, obtained by assuming the matter to be optically thick, which suppresses the directional dependence of the radiation intensity. Details of the derivation may be found in various sources: Mihalas and Mihalas \cite{MM2}, Zel'dovich and Raizer \cite{ZelRai}, Pomraning \cite{Pom}. The gray radiation diffusion equation is a simplification of the MGD equations. It is essentially a one-group equation and is derived by integrating over all frequencies. Surprisingly, it gives very good results in many cases. However, it clearly cannot display frequency-dependent effects. When those are important, it gives incorrect results. Unfortunately, unless one solves a problem with both gray and MGD, one never knows when the former is adequate. We now summarize the paper. Our MGD scheme consists of two parts. Sections~\ref{levelsolve} and \ref{MGanlsys} develop the level-solve algorithm, which is applied on each level. Section~\ref{levelsolve} develops the equations, the discretization, and our $\ptc$ scheme. Section~\ref{MGanlsys} proves three lemmas which determine the initial magnitude of the $\ptc$ parameter $\gs$. % Our philosophy for $\gs$ is as follows. The result of the level solve is the time-advanced radiation group energy density, which physics dictates to be nonnegative. Zeroing anomalously negative values is not an option since they are the correct conservative solution to the linear system that stems from the discretization of the system. Thus, the unphysical result nonetheless conserves energy. The difficulty is avoided if in the original formulation of the linear system $ A x = b$, $A$ is an M-matrix and the right-hand-side (RS) is nonnegative. Since we solve $ A x = b$ using an iterative scheme, the magnitude of $\gs$ is determined to ensure $b \ge 0$, a diagonally dominant $A$, and that the iterations converge. To a large extent, we are guided by Pert~\cite{Pert}, who discusses how and why the solution to a discretization of an equation may be unacceptable from a physical standpoint. For a first reading, section~\ref{MGanlsys} may be skipped; the analysis of the required magnitude of $\gs$ is not needed for the subsequent sections. We note that $\ptc$ is widely used to solve nonlinear systems of equations. It is closely related to the Inexact Newton Backtracking Method by Shahid et al \cite{ShTuWa}. When applying $\ptc$ to a Newton solver, the basic idea is to limit the change to the iterates when one is far from the root but not restrict the change as one approaches the root. With $\ptc$, limiting is done by the magnitude of the pseudo-timestep. Kelley and Keyes \cite{KelKey} put $\ptc$ on a solid analytic framework by examining the three regimes of $\ptc$: small, medium, and large pseudo-timesteps. In the last regime, $\ptc$ recovers Newton's second order of convergence. Our $\ptc$ implementation differs from the norm. Standard applications typically detect when a problem is ``hard'' and then reduce the timestep or some other parameter by an arbitrary amount. However, this method will not work for us because our solver is embedded in a time-dependent multiphysics code with separate modules for compressible gasdynamics, heat conduction, radiation transport. Our MGD solver is called numerous times during the course of a simulation. (If running, with AMR, multiple times per physical time advance.) Although the physical $\dt$ is controlled by various means, and depending on the problem can vary many orders of magnitude, we require a MGD solver that works under all conditions. Our $\ptc$ approach is similar to the one of Shestakov et al \cite{ShHoGr}. We set the initial magnitude of the $\ptc$ parameter to ensure that for the first step, our iteration scheme converges and that the result is physical. We note that our usage of $\ptc$ is nearly equivalent to having the MGD module time-advance not in a single (physical) step $\dt$, but in smaller time increments until the desired time $t^0 + \dt$ is reached. Some colleagues refer to the process as ``sub-cycling'' the radiation module. It is easy to show that the lemmas of Sec.~\ref{MGanlsys} still apply for sub-cycling. Section~\ref{mgamr} describes the second part of our solver, viz., the sync-solve. Section~\ref{rapAMR} contains results. Three problems are presented. The first, in Sec.~\ref{linwin}, displays the accuracy of the method and its convergence properties: first order in time and second order in space. Section~\ref{PTCrobust} demonstrates the utility afforded by $\ptc$. For hard problems, it accelerates convergence; for very hard problems, $\ptc$ is indispensable. Section~\ref{hotball} models the explosive expansion of a hot metal sphere suspended in cold air. The simulation couples all of the code's physics modules. The problem is an ideal candidate for AMR since effects propagate a large distance away from the source, yet in early times, resolution is needed only near the sphere. The problem also demonstrates the necessity of multigroup diffusion. We find that if the sphere's energy is very high, gray diffusion gives the wrong answer. For a 1 MT energy source, our MGD simulation contradicts results of Brode \cite{Bro}, who used gray diffusion. Section~\ref{conclusion} contains concluding remarks. There are three appendices. Appendix~\ref{table} gives a table of exact values for the test problem described in section~\ref{linwin}. Appendix~\ref{apb} discusses situations that may complicate attaining a diagonally dominant matrix when discretizing the multigroup system. Appendix~\ref{apc} presents a spatial convergence analysis of the multigroup system when running in ``production'' mode, that is, with a dominant flux limiter and with AMR. | \label{conclusion} We have described a numerical scheme to solve the radiation multigroup diffusion equations. The scheme is implemented in a radiation-hydrodynamic code with the patch-based AMR methodology, originally proposed by Berger and Oliger \cite{BeOl} for hyperbolic partial differential equations. Our scheme consists of two parts. The first, described in Sections~\ref{levelsolve} and \ref{MGanlsys}, is applied on a {\em level\/} of the AMR grid layout and may be adapted to any code. This part consists of adding $\ptc$ to the ``fully-implicit'' iterative scheme of Axelrod et al \cite{AxDuRh}. $\ptc$ brings an extra degree of robustness and enhances convergence of the Axelrod scheme. We have developed lemmas that determine the minimum magnitude for the $\ptc$ parameter $\tau$ to ensure that the iterations converge and the result is physically meaningful. The appropriate magnitude depends on the problem. Our implementation of $\ptc$ is not optimal---at least for our AMR code architecture. In our code, for each AMR level, we compute a {\em single\/} scalar parameter $\tau$. However, the levels consist of a collection of grids (rectangles in 2D) that need not be connected. If the grids are not connected, they form independent problems. Hence, it would be more efficient to use different $\tau$ for disconnected grids. The second part of our scheme, the sync-solve (SS), addresses a specific need of our code, viz., the requirement of having an energy-conserving result on the composite grid of multiple AMR levels. For the multigroup equations, this part reduces to a coupled system of elliptic equations on the unstructured grid combining all levels. Since the SS is intended to be a small correction to the result of the level solves, we adapted the key element of the ``partial temperature'' scheme of Lund and Wilson \cite{LunWil}. This allowed reducing the multigroup SS to a collection of scalar SS's. We were then able to reuse existing software.% This paper included simulations of three problems. The first two are idealized tests of only the multigroup module. The third is a ``real'' problem, which uses the full capability of the code: AMR, multiple materials, etc. The first problem was chosen because of its non-triviality and the availability of analytic results with which to compare. We obtained excellent agreement and verified the convergence properties of the scheme. The second problem illustrated the benefits brought by $\ptc$. We compared the conventional scheme of Axelrod et al \cite{AxDuRh} with our $\ptc$-modified version. For hard problems, $\ptc$ either decreased run times or ensured convergence in regimes where the conventional scheme diverged. The third problem showed that our multigroup module has been fully integrated into the code and has already extended the scientific frontier. For a high yield air burst at STP, we found that gray diffusion gives an incorrect result during the radiation-dominated regime because gray fails to capture the frequency-dependent effects of the air opacity. | 7 | 10 | 0710.4509 |
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