subfolder
stringclasses 367
values | filename
stringlengths 13
25
| abstract
stringlengths 1
39.9k
| introduction
stringlengths 0
316k
| conclusions
stringlengths 0
229k
| year
int64 0
99
| month
int64 1
12
| arxiv_id
stringlengths 8
25
|
|---|---|---|---|---|---|---|---|
1206
|
1206.4196_arXiv.txt
|
A bouncing cosmology with an initial matter-dominated phase of contraction during which scales which are currently probed with cosmological observations exit the Hubble radius provides a mechanism alternative to inflation for producing a nearly scale-invariant spectrum of cosmological perturbations. In this review article I first discuss the evolution of cosmological fluctuations in the matter bounce scenario, and then discuss various attempts at realizing such a scenario. Observational signatures which will allow the matter bounce to be distinguished from the inflationary paradigm are also discussed.
|
The idea that instead of originating from a Big Bang singularity, the universe has emerged from a cosmological bounce has a long history (see \cite{Novello} for a review with an extensive list of references to the original literature). The group of Professor Novello has made a lot of important contributions to the research on this topic. However, it was realized only fairly recently \cite{Wands, Fabio1} that a bouncing cosmology with a matter-dominated phase of contraction during which scales which are probed today in cosmological observations exit the Hubble radius can provide an alternative to the current inflationary universe paradigm of cosmological structure formation. In this review article, we provide an overview of this ``Matter Bounce" scenario of structure formation, and we discuss some recent efforts at obtaining a non-singular bouncing cosmological background (see also \cite{RHBprev} for reviews comparing the Matter Bounce with inflation and other alternatives to inflation). Inflationary cosmology \cite{Guth} (see also \cite{Brout, Sato, Starob1}) has become the paradigm of early universe cosmology not only because it addresses some of the conceptual problems of Standard Big Bang cosmology such as the horizon, flatness and entropy problems, but because it provided the first causal mechanism for generating the primordial fluctuations which could have developed into the structures we see today on large scales \cite{Mukh} (see also \cite{Press, Sato, Starob2}). More specifically, it predicted a roughly scale-invariant spectrum of cosmological fluctuations which in simple models of inflation have a slight red tilt and which are Gaussian and nearly adiabatic. These predictions were spectacularly confirmed in recent precision observations of Cosmic Microwave Background (CMB) anisotropies \cite{WMAP}. On the other hand, it was known since long before the development of inflationary cosmology that any model which produces a roughly scale-invariant and almost adiabatic spectrum of cosmological perturbations will be a good match to observations \cite{Peebles, Sunyaev}. Inflationary cosmology is simply the first model based on fundamental physics which yielded such a spectrum. In the mean time, other models have been developed which predict this kind of spectrum, e.g. the Ekpyrotic universe \cite{Ekp}, ``string gas cosmology" \cite{SGC}, the ``Varying Speed of Light" proposal \cite{Moffat}, the ``Conformal Universe" \cite{Rubakov}, and the Matter Bounce scenario. In the following section, I shall show how quantum vacuum fluctuations originating on sub-Hubble scales and exiting the Hubble radius in a matter-dominated phase of contracting develop into a scale-invariant spectrum of curvature perturbations. As is well known since the pioneering work of Hawking and Penrose (see e.g. \cite{HE} for a textbook description)), a cosmological singularity is unavoidable if space-time is described in terms of General Relativity (GR) and if matter obeys certain energy conditions. Thus, in order to obtain a bouncing cosmology it is necessary to either go beyond Einstein gravity, or else to introduce new forms of matter which violate the key energy conditions. In Section 3 of this review I will describe some concrete and recent models which yield a bouncing cosmology (the reader is referred to \cite{Novello} for an overview of early work on obtaining bouncing universes).
|
The ``Matter Bounce" is an alternative to cosmological inflation as a mechanism for generating an almost scale-invariant spectrum of cosmological fluctuations. As in the case of inflationary cosmology, the spectrum has a slight red tilt since smaller scales exit the Hubble radius when the radiation component of matter is more important and the vacuum slope of the spectrum (which corresponds to $n_s = 3$) is rearing its head. In fact, there is a transition to an $n_s = 3$ spectrum on small scales which exit the Hubble radius in the radiation phase of contraction. The current observational constraints on a spectrum with such a transition from $n_s = 1$ on large scales to $n_s = 3$ on small scales were studied in \cite{LiHong}, with the result that the power spectrum of quasar absorption lines already mildly constrains a model with no entropy production during the bounce. If there is entropy production during the bounce, then the radiation phase of contraction is shorter than the radiation phase of expansion and the observational constraints recede. The spectrum of gravitational radiation produced in a matter bounce is also scale-invariant. At first glance, there will be less enhancement of the amplitude during the bounce phase than for scalar metric fluctuations, and hence the predicted tensor to scalar ratio $r$ will be consistent with current constraints. However, a detailed study of this issue is needed. Since the curvature fluctuation ${\cal R}$ increases on super-Hubble scales during the contracting period, terms in the general expression for the bispectrum (three point function of the curvature fluctuation variable) derived in \cite{Malda} will be important, leading to a shape of the bispectrum which is different from the one obtained in simple inflationary models \cite{bispectrum}. Since the analog of the inflationary slow-roll parameter is of order unity in the case of the Matter Bounce, the predicted amplitude of the bispectrum will be of order unity \cite{bispectrum}. Thus, amplitude and shape of the bispectrum provide a way to differentiate between the Matter Bounce and simple inflationary models. Possibly the most serious problem which the Matter Bounce scenario faces is the anisotropy problem (see e.g. \cite{BingKan}). The only solution of this problem which we know of now is to postulate a phase of Ekpyrotic contraction following the initial matter contraction phase. In the model of \cite{Cai3} this solves the anisotropy problem, whereas in the original New Ekpyrotic scenario the anisotropy problem may re-appear during the bounce phase \cite{BingKan}. I have focused on single bounce models. Cyclic bouncing background cosmologies face several challenges. In addition to the ``heat death" problem (entropy increasing from cycle to cycle), there is the problem that the dynamics of cosmological fluctuations breaks the cyclicity of the model since fluctuations grow on super-Hubble scales during the periods of contraction and lead to a jump in the index $n_s$ of the power spectrum by $\delta n_s = -2$ from cycle to cycle \cite{processing}. Thus, the model loses predictability \footnote{Note that the cyclic version of the Ekpyotic scenario \cite{Ekpcyclic} does not face this problem since the evolution of the scale factor of our universe is not cyclic in the model, only the separation between the branes in the higher-dimensional framework of the model.}.
| 12
| 6
|
1206.4196
|
1206
|
1206.6652_arXiv.txt
|
Since the Standard Model most probably cannot explain the large value of CP asymmetries recently observed in $D$-meson decays we propose the fourth quark-lepton generation explanation of it. As a byproduct weakly mixed leptons of the fourth generation make it possible to save the baryon number of the Universe from erasure by sphalerons. An impact of the 4th generation on BBN is briefly discussed.
|
Recently LHCb collaboration has measured the unexpectedly large CP violating asymmetries in $D\to\pi^+ \pi^-$ and $D\to K^+ K^-$ decays \cite{1}: \begin{equation} \Delta A_{CP}^{LHCb} \equiv A_{CP}(K^+ K^-) - A_{CP}(\pi^+ \pi^-) = [-0.82 \pm 0.21 (\rm{stat.}) \pm 0.11 (\rm{syst.})]\% \;\; , \label{1} \end{equation} where \begin{equation} A_{CP}(\pi^+ \pi^-) = \frac{\Gamma(D^0 \to \pi^+ \pi^-) -\Gamma(\bar D^0 \to \pi^+ \pi^-)}{\Gamma(D^0 \to \pi^+ \pi^-) +\Gamma(\bar D^0 \to \pi^+ \pi^-)} \label{2} \end{equation} and $A_{CP}(K^+ K^-)$ is defined analogously. This result was later confirmed by CDF collaboration, which obtained~\cite{2}: \begin{equation} \Delta A_{CP}^{CDF} = [-0.62 \pm 0.21 (\rm{stat.}) \pm 0.10 (\rm{syst.})]\% \;\; . \label{3} \end{equation} The most important question concerning experimental results (\ref{1}) and (\ref{3}) is whether in the Standard Model the CP-violation (CPV) in these decays can be as large as 0.5\% - 1\%. In the Standard Model the CPV in $D(\bar D) \to\pi^+ \pi^-$ decays originates from the interference of the tree and penguin diagrams shown in Fig. 1. For $D(\bar D)\to K^+ K^-$ decays $d$-quarks in these diagrams should be substituted by $s$-quarks. \begin{figure} \begin{center} \includegraphics[scale=0.6]{tree1.eps} \caption{Quark diagrams describing $D\longrightarrow\pi^+ \pi^-$ decay in the Standard Model. A wavy line denotes $W$-boson, a curly line -- gluon.} \end{center} \end{figure} It is convenient to present the penguin diagram contribution to $D\to\pi^+ \pi^-$ decay amplitude in the following form \cite{222}: \begin{eqnarray} V_{cd}V_{ud}^* f(m_d) + V_{cs}V_{us}^* f(m_s) + V_{cb} V_{ub}^* f(m_b) = \nonumber \\ = V_{cd} V_{ud}^*[f(m_d)-f(m_s)] + V_{cb} V_{ub}^* [f(m_b) - f(m_s)] \;\; , \label{4} \end{eqnarray} attributing the first term to the tree amplitude and considering the second term only as the penguin amplitude. In the case of $D\to K^+ K^-$ decay the following presentation is useful \cite{222}: \begin{eqnarray} V_{cd}V_{ud}^* f(m_d) + V_{cs}V_{us}^* f(m_s) + V_{cb} V_{ub}^* f(m_b) = \nonumber \\ = V_{cs} V_{us}^*[f(m_s)-f(m_d)] + V_{cb} V_{ub}^* [f(m_b) - f(m_d)] \, , \label{5} \end{eqnarray} where the first term is attributed to the tree amplitude while the second one is the penguin amplitude. Denoting the absolute values of $D\to\pi^+\pi^-$ decay amplitudes by $T$ and $P$ we get: \begin{eqnarray} A_{\pi^+\pi^-} & = & T\left[1+\frac{P}{T} e^{i(\delta-\gamma)}\right] \;\; , \nonumber \\ \bar A_{\pi^+\pi^-} & = & T\left[1+\frac{P}{T} e^{i(\delta+\gamma)}\right] \;\; , \label{6} \end{eqnarray} where $\delta$ stands for the difference of the strong interaction phases of the tree and the penguin amplitudes, while $\gamma\approx 70^0$ is the phase of $V_{ub}$ (the product $V_{cd} V_{ud}^*$ as well as $V_{cb}$ are practically real in the standard parametrization of the CKM matrix). From eq.~(\ref{6}) for the CPV asymmetry we obtain: \begin{equation} A_{CP}(\pi^+ \pi^-) = 2\frac{P}{T} \sin\delta \sin\gamma \;\; , \label{7} \end{equation} where in the denominator of (\ref{2}) we neglect the terms of the order of $P/T$ and $(P/T)^2$ which is a very good approximation because $P/T\sim |V_{cb}V_{ub}^*|/V_{cd} \ll 1$. Here $\sin\gamma$ is close to unity and we use this value in what follows. Let us present an argument demonstrating that $\delta$ can also be close to 90$^0$. The tree diagram gives dominant contribution to the $D\to\pi\pi$ decay rates. The corresponding to it 4-fermion Hamiltonian has parts with isospin 1/2 and 3/2. That is why the produced $\pi$-meson may have isospin zero or two. So three decay probabilities, $D^+ \to \pi^+ \pi^0$, $D^0\to\pi^+ \pi^-$, and $D^0 \to \pi^0\pi^0$, depend on the absolute values of the decay amplitudes $A_0$ and $A_2$ and their strong phases difference $\delta_0 -\delta_2$. From the experimentally measured branching ratios \cite{3}: $$ {\rm Br}(D^+ \to\pi^+ \pi^0) = [12.6\pm 0.9]\cdot 10^{-4} \; , \;\; {\rm Br}(D^0 \to\pi^0 \pi^0) = [8.0\pm 0.8]\cdot 10^{-4} \; , $$ \begin{equation} {\rm Br}(D^0 \to\pi^+ \pi^-) = [13.97\pm 0.26]\cdot 10^{-4} \label{8} \end{equation} we find for the phase difference of the amplitudes with $I=0$ and $I=2$: \begin{equation} |\delta_0 -\delta_2| = 86^0 \pm 4^0 \;\; . \label{9} \end{equation} In eq. (\ref{7}) $\delta$ stands for the difference of the strong phases of penguin amplitude which has $I=1/2$ and produces pions with $I=0$ and tree amplitude, which has parts with $I=1/2$ and $I=3/2$ and produces pions with $I=0$ and $I=2$, that is why $\delta \neq \delta_0 -\delta_2$. Nevertheless eq. (\ref{9}) demonstrates that $\delta$ can be large, and so we substitute $\sin\delta =1$ into eq. (\ref{7}). In the limit of $U$-spin ($d\leftrightarrow s$ interchange) symmetry the tree amplitude of $D(\bar D)\to K^+ K^-$ decay differs by sign from that of $D(\bar D) \to\pi^+ \pi^-$ decay, while the penguin amplitudes of these decays are equal, that is why \begin{equation} A_{CP}(K^+ K^-) = -A_{CP}(\pi^+\pi^-) \;\; . \label{10} \end{equation} However since \cite{3} \begin{equation} {\rm Br}(D^0 \to K^+ K^-) = [39.4\pm 0.7] \cdot 10^{-4} \label{11}\,, \end{equation} we obtain from eq.~(\ref{8}) that $|A_{K^+ K^-}/A_{\pi^+\pi^-}| \simeq 1.7$ and $U$-spin symmetry is heavily broken in $D$ decays. Nevertheless let us suppose that (\ref{10}) is not badly violated, so finally we get: \begin{equation} \Delta A_{CP} = 4\,\frac{P}{T} \label{12}\,. \end{equation} Now let us try to understand if in the Standard Model we can obtain \begin{equation} \frac{P}{T} = 1.8 \cdot 10^{-3} \label{13}\,, \end{equation} which is needed to reproduce the average value of the LHCb and CDF results.
|
In Introduction we determined what ratio of the penguin to the tree amplitudes of $D\to\pi^+ \pi^-$ decay is needed to get the observed CP asymmetry. In Section 2 we found that the factorization describes the tree amplitude with good accuracy; concerning the penguin amplitude it appears to be twenty times smaller than one needs to describe the experimental data on $A_{CP}$. In Section 3 we demonstrated that in the case of $B\to \pi^+ K^0$ decay the factorization underestimates the penguin amplitude by factor 2. In the case of $K_S \to \pi^+ \pi^-$ decay the penguin amplitude is enhanced by factor 2-3 in comparison with the factorization result. Thus if confirmed on larger statistics and future systematics result (\ref{1}) demands New Physics. In Section 5 we demonstrated that the fourth quark-lepton generation may enhance the penguin amplitude describing the experimental data. If the leptons of the fourth generation weakly mix with three light generation leptons, then the baryonic charge generated at high scale escapes the erasure by sphalerons and survives till now according to the results presented in Section 6. We are grateful to S.I. Blinnikov for the illuminating discussion on the chemical potentials, to V.A. Rubakov for the clarifying discussion on the baryon density in the unbroken electroweak phase, and to J. Zupan for the remark concerning $D\rightarrow K^{0}\bar {K^{0}}$ decay. A.D., S.G., and M.V. acknowledge the support of the grant of the Russian Federation government 11.G34.31.0047. S.G. and M.V. are partially supported by the grants RFBR 11-02-00441, 12-02-00193 and by the grant NSh-3172.2012.2. S.G. is partially supported by the grant RFBR 10-02-01398. \bigskip \begin{center} {\large \bf Appendix} \end{center} \setcounter{equation}{0} \def\theequation{A.\arabic{equation}} \vspace{3mm} Below we derive equations used in Section 6 to find the dependence of the baryon asymmetry of the Universe on the sphaleron freeze-out temperature. In this Appendix we closely follow paper~\cite{11}. Being interested in the values of the asymmetries at sphaleron freeze-out temperature we should assume that the electroweak phase transition already has occured and the neutral Higgs boson condenses. That is why the Higgs boson chemical potential is zero. Sometimes in the literature the baryon density in the electroweak unbroken phase is looked for. In this case the Higgs boson does not condense and its chemical potential is nonzero. To find it an additional equation is needed. It is provided by the condition that the density of charges with which the massless bosons interact should be zero, and in an unbroken phase there are two such charges: the hypercharge and the third projection of a weak isospin. The baryon density in the unbroken phase is analyzed, for example, in book \cite{15} and it differs from its value in a broken phase. Since the right-handed components of quarks and leptons emitting neutral Higgs transform to the left-handed components the chemical potential of both components are equal: $\mu_{u_R} = \mu_{u_L} \equiv \mu_u$, $\mu_{d_R} = \mu_{d_L} \equiv \mu_d$, $\mu_{e_R} = \mu_{e_L} \equiv \mu_e$. The analogous relations are valid for the particles of the second and third families. The right-handed neutrinos of three light generations are not thermalized and should not be taken into account (see the end of Sect. 6). The fourth generation right-handed neutrinos, being heavy, rapidly thermalize: $\mu_{N_R} = \mu_{N_L} \equiv \mu_N$. The chemical potentials of up and down weak isospin components are related by $W^-$ chemical potential: $\mu_d = \mu_W + \mu_u$, $\mu_e = \mu_W + \mu_\nu$, $\mu_E = \mu_W + \mu_N$. Mixing of quarks of four families and leptons of three families equilibrates the chemical potentials of the particles with the identical gauge quantum numbers. As a result four independent chemical potentials remain: $\mu_u$, $\mu_N$, $\mu_W$ and $\mu\equiv \mu_{\nu_1} + \mu_{\nu_2} + \mu_{\nu_3} \equiv 3\mu_\nu$. The particle number densities depend on their (Fermi or Bose) statistics, temperature, chemical potential, and masses. The chemical potential of an antiparticle is opposite to that of the particle. The asymmetries and, hence, chemical potentials are very small. Expanding the equilibrium integrals for the asymmetry over $\mu$ we obtain: $$ n_p = \frac{g_p}{\pi^2}T^3\left(\frac{\mu}{T}\right) \int\limits_x^\infty y\sqrt{y^2 - x^2} \frac{e^y}{(1\pm e^y)^2}dy = $$ \begin{equation} = \left\{ \begin{array}{ll} \frac{g_p T^3}{3} \left(\frac{\mu}{T}\right) \alpha_b(x) \; , & {\rm if} ~ p ~ {\rm is ~ a ~ boson} \\ \frac{g_p T^3}{6} \left(\frac{\mu}{T}\right) \alpha_f(x) \; , & {\rm if} ~ p ~ {\rm is ~ a ~ fermion} \;\; , \label{A.1} \end{array} \right. \end{equation} where $g_p$ is the number of the degrees of freedom of the particle $p$ ($g_q = g_l =2$, $g_\nu =1$, $g_N =2$, $g_W =3$) and $x=m/T$. Functions $\alpha(x)$ are normalized in such a way that $\alpha_b(0)=\alpha_f(0) =1$. In what follows we take into account the nonzero masses of the particles of the fourth generation, of $t$-quark, and of $W$-boson. The condition of electroneutrality of the primeval plasma looks as: \begin{eqnarray} Q & = & 3\frac{2}{3}[2(\alpha_u + \alpha_c +\alpha_t +\alpha_{t'})\mu_u] - 3\,\frac{1}{3}[2(\alpha_d + \alpha_s + \alpha_b +\alpha_{b'}) \times \nonumber \\ &\times &(\mu_W + \mu_u)] -2[(\alpha_e + \alpha_\mu + \alpha_\tau)(\mu_W + \mu_\nu)] - 2\alpha_E(\mu_W + \mu_N) - \nonumber \\ & - & 3\cdot 2 \alpha_W \mu_W = 0 \;\; , \label{A.2} \end{eqnarray} \begin{equation} (1+2\alpha_t + 2\alpha_{t'} -\alpha_{b'})\mu_u - (6+\alpha_{b'} +\alpha_E + 3\alpha_W)\mu_W -\mu - \alpha_E \mu_N = 0. \label{A.3} \end{equation} Here and below we omit irrelevant factor $T^{2}/6$. The sphaleron transition converts $qqql$ combination of each generation into vacuum, which gives: \begin{equation} 12\mu_u + 8\mu_W + \mu + \mu_N =0 \;\; . \label{A.4} \end{equation} The remaining two equations are two superpositions of $B^\prime$, $L$, and $L_4$ conserved under sphaleron transitions thus being equal to their initial values. The expressions for these quantities look like: \begin{equation} L_4 = 2\alpha_E \mu_E + 2\alpha_N \mu_N = 2(\alpha_E + \alpha_N) \mu_N + 2\alpha_E \mu_W \;\; , \label{A.5} \end{equation} \begin{equation} L = 2(\alpha_e + \alpha_\mu + \alpha_\tau)\mu_e +(\alpha_{\nu_e}+\alpha_{\nu_\mu} + \alpha_{\nu_\tau}) \frac{\mu}{3} = 3\mu + 6\mu_W \;\; , \label{A.6} \end{equation} \begin{eqnarray} B^\prime & = & 2\cdot 3 \cdot \frac{1}{3}[(\alpha_u + \alpha_c + \alpha_t +\alpha_{t'})\mu_u +(\alpha_d + \alpha_s + \alpha_b +\alpha_{b'})\mu_d] = \nonumber \\ & = & 2(2+\alpha_t +\alpha_{t'})\mu_u + 2(3+\alpha_{b'})(\mu_u + \mu_W) \;\; . \label{A.7} \end{eqnarray} Thus we have four equations which determine the chemical potentials: (\ref{A.3}), (\ref{A.4}), and the remaining two: \begin{eqnarray} B^\prime -L -L_4 &=&2(5+\alpha_t +\alpha_{t'} +\alpha_{b'})\mu_u +2(\alpha_{b'} -\alpha_E)\mu_W -\nonumber\\ &&\hspace{2.55cm}-3\mu -2(\alpha_E +\alpha_N)\mu_N = 0 \;\; , \label{A.8}\\ L - 3L_4 &=& 6(1-\alpha_E)\mu_W + 3\mu -6(\alpha_E +\alpha_N)\mu_N = 3\Delta \;\; , \end{eqnarray} where we take the initial values analogous to those of ref.~\cite{11}: $B_0^\prime = L_0 = 3\Delta$ and $L_4^0 =0$. When temperature is much larger than the masses of all the particles, all $\alpha_i$ are equal to one we obtain: \begin{equation} \frac{B^\prime}{\Delta}\left|_{T\gg m_i} = -\frac{11}{179} \right. \;\; . \label{A.9} \end{equation} If the right-handed neutrinos of three light generations thermalized then the equation (\ref{A.6}) would be substituted by \begin{equation} L=4\mu +6\mu_W \;\; , \label{A.10} \end{equation} and the baryon asymmetry at $T\gg m_i$ would vanish.
| 12
| 6
|
1206.6652
|
1206
|
1206.3077_arXiv.txt
|
In this paper we study gravitational lensing in the strong field limit from the perspective of cosmic censorship, to investigate whether or not naked singularities, if at all they exist in nature, can be distinguished from black holes. The spacetime which we explore from this perspective is JMN metric which represents a spherically symmetric solution to the Einstein field equations with anisotropic pressure and contains a naked singularity at the center. JMN geometry is matched with the \sch metric to the exterior at a finite radius. This metric was recently shown to be a possible end state of gravitational collapse of a fluid with zero radial pressure. In the presence of the photon sphere gravitational lensing signature of this spacetime is identical to that of \sch black hole with infinitely many relativistic images and Einstein rings, all of them located beyond a certain critical angle from optic axis and the inner relativistic images all clumped together. However, in the absence of the photon sphere, which is the case for a wide range of parameter values in this spacetime, we show that we get finitely many relativistic images and Einstein rings spaced reasonably apart from one another, some of which can be formed inside the critical angle for the corresponding \sch black hole. This study suggests that the observation of relativistic images and rings might, in principle, allow us to unravel the existence of the naked singularity in the absence of the photon sphere. Also the results obtained here are in contrast with the earlier investigation on JNW naked singularities where it was shown that the radial caustic is always present in the absence the photon sphere, which is not the case with JMN geometry where radial caustic is absent. We also point out the practical difficulties that might be encountered in the observation of the relativistic images and suggest that new dedicated experiments and techniques have be developed in future for this purpose.
|
Deflection of light by massive bodies and therefore the phenomenon of gravitational lensing is a prediction of \gr which has helped in observationally testing Einstein theory against Newtonian gravity. In fact it was one of the first successfully verified predictions of \gr in 1919. Since then there has been numerous studies of gravitational lensing, both theoretically and observationally. But for good reasons, mostly these have been confined to the weak field approximation. The last decade, however, has seen a great rise in the interest in \grle in the strong field regime. This is important as a test of \gr in itself (because almost all generalizations of \gr should by construction reproduce the same weak field limit), as well as for the testing of various compact object scenarios in \gr (because they too have the same weak field). In particular, it is important to study whether and how strong field lensing can distinguish between various end states of gravitational collapse of massive matter clouds (such as a massive star continually collapsing at the end of its life cycle), viz. black hole and naked singularity. This is also important from the perspective of the cosmic censorship conjecture. Cosmic censorship conjecture was proposed by Penrose in order to get rid of the naked singularities in the real world around us \cite{pen69}. However the cosmic censorship conjecture is not yet proved even several decades after it was put forward. There were many studies recently where it was shown that the black holes as well as naked singularities are formed in a continual gravitational collapse of a matter cloud of reasonable matter field starting from a regular initial data \cite{Joshi1,Joshi2}. Thus naked singularities might occur in nature. Their occurrence or otherwise is hard to infer from purely theoretical investigations as it is extremely difficult to solve the Einstein equations in an astrophysically realistic scenario. Thus one could take a phenomenological approach, where the consequences of the occurrence of the naked singularities computed theoretically are compared with the observations to either confirm or rule out the existence of the naked singularities. In this paper we explore the gravitational lensing from such a perspective. We note here that the strong gravitational lensing in JNW spacetime \cite{vnc98,nsl.ve}, lensing in post-Newtonian framework for Kerr geometry \cite{wp} as well as for the rotating generalization of JNW spacetime \cite{gy} and the investigation of the shape and the position of the shadow in Kerr and Tamimatsu-Sato spacetimes \cite{bambi1,bambi2,maeda} has been done recently to address the same question. Early works on strong field lensing were by Darwin \cite{dar59,dar61}, who studied the behavior of null geodesics in strong field regime of \sch black hole and pointed out the divergence of \eda as the distance of closest approach of the geodesics approaches photon sphere. Strong field lensing with a lens equation was studied by Virbhadra and Ellis \cite{sbhl.ve}, who examined strong field lensing in Schwarzschild black holes and showed that there could in principle be infinite relativistic images on each side of the black hole when a light ray with small enough impact parameter ( distance of closest approach close enough to photon sphere) goes around one or several times around the black hole before reaching the observer. Earlier, lens equation for spherically symmetric static spacetimes that goes beyond the weak field small-angle approximation was studied by Virbhadra ,Narasimha and Chitre in \cite{vnc98}. The Virbhadra-Ellis type lens equation has also been applied to boson star by D\c{a}browski and Schunck \cite{das00}, to a fermion star by Bili{\'c}, Nikoli{\'c} and Viollier \cite{bnv00}. As one of the first steps towards using strong field lensing to probe the cosmic censorship question, Virbhadra and Ellis have used this lens equation to study and compare gravitational lensing by normal black holes and by naked singularities modeled by the Janis, Newman, Winicour metric (JNW solution)\ct{nsl.ve}. It is worthwhile to extend this line of work to other novel, more interesting and if possible more realistic naked singularity models. With this in mind we consider here the class of solutions recently obtained by Joshi, Malafarina and Narayan \ct{jmn} as end state of certain dynamical collapse scenarios in a toy example. JMN metric is a solution of Einstein field equations with an anisotropic pressure fluid and has a naked singularity at the center. It is matched to the \sch metric at a certain radius. We refer to it here as JMN naked singularity from now on. It is worthwhile to mention that, not only the presence of the central naked singularity but also the value of the radius at which the interior solution is matched to exterior \sch geometry plays a crucial role in determining gravitational lensing observables. We should also mention that exact lens equations were proposed by \cite{fkn} for arbitrary spacetime and also by \ct{per04} for spherically symmetric case. Bozza et al. have defined and analytically calculated strong field limit observables in spherically symmetric spacetimes endowed with a photon sphere \cite{boz02,bcis01}. In such a situation strong lensing from various alternatives/modifications of \sch geometry in modeling the galactic center has been studied. For example lensing from regular black holes was studied in \cite{eis11} and lensing from stringy black holes was studied in \cite{bha03}. However the basic qualitative features in a lensing scenario in the presence of a photon sphere is very similar to \sch case and is ineffective in probing the geometry beyond the photon sphere. Strong field lensing would be much easily able to probe differences from \sch spacetime if geometry being studied will be without a photon sphere. As we will see for the family of solutions studied in this paper, when the geometry has a photon sphere the lensing signatures are exactly identical to \sch black hole case while in the absence of photon sphere it is greatly different. In this work, the galactic supermassive compact object is analyzed as a strong gravity lens to illustrate these characteristics. This paper is organized as follows. In section \ref{bf} we introduce the basic formalism in brief. In section \ref{galmod} we discuss the lens model with galactic supermassive dark object as the lens and in section \ref{sch} we discuss the lensing signatures when it is modeled as a \sch black hole. We discuss the naked singularity spacetime we intend to study and lensing in this background in \ref{jnm} and compare this with \sch back hole and JNW solution in \ref{schcomp} \& \ref{jnwcomp} respectively. In section \ref{obs}, we discuss the implications of going beyond point source approximation for our study and in \ref{binary} we briefly discuss how binary systems could be useful for probing question of cosmic censorship via gravitational lensing. Finally, we discuss the main results and conclude with a general discussion in section \ref{rem}.
|
\label{rem} In this paper we studied the strong gravitational lensing from the perspective of cosmic censorship and explored the possibility of distinguishing black holes from naked singularities. We modeled the galactic central supermassive dark object initially by a black hole and then by naked singularity. We studied the gravitational lensing of the source in a near aligned configuration at a distance from a galactic center approximately comparable to distance of the sun from the center. The \sch black hole has a photon sphere. Thus apart from a pair of nonrelativistic images and a nonrelativistic Einstein ring, infinitely many relativistic images and Einstein rings clumped together. No images and Einstein rings lie in the region between the optic axis and $\theta=16.8$ \mas. Also all the images that are clumped together are highly demagnified as compared to the first relativistic image with a small separation between them of the order of $0.1$ \mas. We then model the galactic center object as JMN solution which was recently shown to occur as an end state of the gravitational collapse of a fluid with zero radial pressure but non-vanishing tangential pressure. This solution has two parameters, namely mass and another parameter $M_0$. The spacetime is divided into two parts. Exterior metric is \sch spacetime with same mass as that of the \sch black hole considered earlier. Interior metric contains a central naked singularity with the boundary located at the radius $R_b=\frac{2M}{M_0}$.The two metrics are connected across the boundary by $C^2$ matching. In the parameter range $M_0 \ge \frac{2}{3}$ the \sch photon sphere is present in the geometry and the gravitatioanl lensing signature of JMN spacetime is identical to that of the \sch black hole. When $M_0 \le 0.475$, the photon sphere is absent. But no relativitic bending of light and thus no relativistic images possible. This behavior is different from the \sch black hole. The interesting parameter range is when $0.475<M_0 <\frac{2}{3}$. The photon sphere is absent. But the relativistic images and Einstein rings can form and their number increases with increasing value of the parameter $M_0$. The images and rings are well separated from one another and happen to lie in the forbidden region for \sch black hole, within a distance from the optic axis of $\theta=16.8$ \mas. Their magnification is also comparable. Thus the strong gravitational lensing signature is qualitatively different from \sch black hole. The gravitational lensing in the absence of the photon sphere is qualitatively different in JMN and JNW spacetimes. In both the geometries relativistic images are present in an appropriate parameter range. However, there are no radial caustics in the JMN geometry, while radial caustic is always present in the JNW spacetime in the absence of the photon sphere. However, there are practical difficulties as far as observation of relativistic images and rings are concerned with the telescopes and techniques currently being used. We require \mas resolution which can be achieved with VLBI. However magnification of the images which is of the order of $\mu=10^{-22}$ is too small. Relativistic Einstein rings formed due to the lensing of the star with the size comparable to sun, will be $10^{17}$ times weaker as compared to the nonrelativistic \sch Einstein ring and thus will not be seen since the current instruments do not have dynamical range over seventeen orders of magnitude of brightness. Keeping this in mind, new techniques and instruments must be developed in the future which will be able to observe the Einstein rings and can unravel the nature of the galactic central supermassive object. We also suggest that the appearance and disappearance of the outer Einstein ring as the source crosses diamond shaped caustic more than once can possibly shed light on the possible existence of the naked singularity in the binary system of a naked singularity and a massive star. We wish to explore this situation in the future. In this paper we studied a naked singularity geometry arising out of a toy calculation of dynamical collapse of a matter with only the tangential pressure. It would be interesting to study more realistic cases e.g. with the inclusion of the radial pressure.
| 12
| 6
|
1206.3077
|
1206
|
1206.3153.txt
|
Discovered just 20 years ago \citep{1993Natur.362..730J}, the Kuiper Belt holds a vast population of icy bodies orbiting the Sun beyond Neptune. Stored at very low temperatures ($\sim$30-50~K), the Kuiper Belt Objects (KBOs) are expected to be well-preserved fossil remnants of the solar system formation. Presently, $\sim$1600 KBOs have been identified and classified into several dynamical families \citep[see Appendix \ref{app:database} and][for a review]{2008ssbn.book...43G}. KBOs which dynamically evolve to become Jupiter Family Comets (JFCs) form a transient population, the Centaurs, with short-lived chaotic orbits between Jupiter and Neptune \citep{Kow77,1980MNRAS.192..481F,1997Icar..127...13L}. Between 1998 and 2003, we witnessed a debate on the surface colors of KBOs and Centaurs. One team used very accurate surface colors and detected that KBOs were separated into two distinct color groups \citep{1998Natur.392...49T,2000Natur.407..979T,2003Icar..161..181T}. Other teams did not find evidence for such color bimodality \citep{1999Icar..142..476B,2001AJ....122.2099J, 2002A&A...389..641H}. Careful reanalysis of the data by \cite{2003A&A...410L..29P} indicated that only the Centaurs display bimodal colors, \ie they are distributed in two distinct color groups, one with neutral solar-like colors, and one with very red colors. KBOs on the other hand exhibit a broad continuous color distribution, from neutral to very red, with no statistical evidence for a color gap between the extrema \citep[][for a review]{2008ssbn.book..105T}. The relevance of this controversy lays on two possible interpretations: i) KBOs and Centaurs are composed of intrinsically different objects, with distinct compositions, which probably formed at different locations of the protosolar disk, ii) KBOs and Centaurs are originally similar but evolutionary processes altered them differently, hence their color diversity. Most research focused on the latter hypothesis, offering little improvement on our understanding of the color distributions. \cite{1996AJ....112.2310L} proposed that the competition between a reddening effect of irradiation of surface ices \citep{1987JGR....9214933T} and a bluing effect due to collisional induced resurfacing of fresh, non-irradiated, ices might generate the observed surface colors. The same authors, however, rejected this model as being the primary cause of the color diversity, due to the lack of predicted rotational color variations \citep{2001AJ....122.2099J}. Based on the same processes, \cite{2002P&SS...50...57G} proposed a more complex treatment of the irradiation process, by implying an intricate structure of differently irradiated subsurface layers. However, the collisional resurfacing effects became very hard to model, thus making it very hard to provide testable predictions. Later, \cite{2003Icar..162...27T} showed that the collisional energies involved in different parts of the Kuiper Belt did not corroborate the possible link between surface colors and non-disruptive collisions. \cite{2004A&A...417.1145D} refined the first-mentioned model by considering the effects of a possible cometary activity triggered by collisions, and a size/gravity-dependent resurfacing. Cometary activity can modify the surface properties through the creation of a neutral-color dust mantle. \cite{2002AJ....123.1039J} suggested that this process could explain why no JFCs are found with the ultra-red surfaces seen in about half of the Centaurs. It has also been proposed that the sublimation loss of surface ice from a mixture with red materials may be sufficient to make the red material undetectable in the visible wavelengths \citep{2009Icar..199..560G}. These might explain the Centaur color bimodality, as long as all were red when migrating inwards from the Kuiper Belt. Although promising, these models did not provide an explanation for the color bimodality of Centaurs, as they fail to reproduce the bluest colors observed and their frequency. %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%
|
In this work we analyze the $B-R$ color distribution as a function of $H_R(\alpha)$ magnitude for 253 Centaurs and KBOs, including 10 new measurements, and with the information on their NIR spectral features. Using the known diameters, $D$, and albedos, $p_V$, of 74 of these objects we verify that $H_R$ and $D$ correlate very strongly ($\rho=-0.92^{+0.03}_{-0.02}$, $SL\ll 0.01\%$) validating $H_R$ as a good proxy for size. Further, through simulations, we show that not correcting $H_R(\alpha)$ to $H_R(\alpha=0^{\circ})$ does not change any of the global results. Our analysis shows: \begin{enumerate} \item The $B-R$ \vs $H_R(\alpha)$ color distribution is $\mathcal{N}$-shaped, evidencing that $B-R$ colors are probably dominated by a size effect independent from dynamical classification. \item Small objects, including both KBOs and Centaurs, display a bimodal structure of $B-R$ colors at $0.1\%$ significance level (\ie objects with $H_R(\alpha)\geqslant 6.8$, or $D_{km}\lesssim 165$, assuming $p_R=0.09$) with the `gap' centered at $B-R\sim1.60$. Removing Centaurs from the sample reduces greatly the sampling on small objects reducing also the significance of the result to $3.8\%$. \item Large objects evidence also for a bimodal structure, with minimum significance of $0.9\%$, for $H_R(\alpha)\lesssim5.0$ ($D_{km}\gtrsim 380$, assuming $p_R=0.09$), and color `gap' centered at $B-R\sim1.25$. Reasonable evidence for this bimodality starts when considering only objects with $H_R(\alpha)\lesssim5.6$ ($D_{km}\gtrsim 290$) dropping below the critical $5\%$ when reaching $H_R(\alpha)\lesssim4.4$ ($D_{km}\gtrsim 500$). However, this behavior seems dominated by the presence of 7 Haumea collisional family objects which `cluster' at the lower left leg of the $\mathcal{N}$-shape. Once removed, there is no statistical evidence against compatibility with a random unimodal distribution for the larger KBOs. \item Intermediate sized objects do not show incompatibility with a continuum of $B-R$ colors (\ie $6.8>H_R(\alpha)>5.0$, or $165\lesssim D_{km} \lesssim 380$, assuming $p_R=0.09$). These objects seem too large to be remnants from disruptive collisions and too small to hold cryovolcanic activity. They might be the best targets to study the combined effects of different birthplaces, different space weathering, and different thermal processing. Further studies are encouraged. \item Inspecting the NIR spectral properties against $B-R$ colors shows no obvious link between the colors and the chemical composition of the objects' surfaces. \end{enumerate} %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%% %%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%%
| 12
| 6
|
1206.3153
|
|
1206
|
1206.3307_arXiv.txt
|
We present the discovery of a 0$\farcs$2 companion to the young M dwarf GJ 3629 as part of our high contrast adaptive optics imaging search for giant planets around low-mass stars with the Keck-II and Subaru telescopes. Two epochs of imaging confirm the pair is co-moving and reveal signs of orbital motion. The primary exhibits saturated X-ray emission, which together with its UV photometry from $GALEX$ point to an age younger than $\sim$300~Myr. At these ages the companion lies below the hydrogen burning limit with a model-dependent mass of 46~$\pm$~16~\Mjup \ based on the system's photometric distance of 22~$\pm$~3~pc. Resolved $YJHK$ photometry of the pair indicates a spectral type of M7~$\pm$~2 for GJ~3629~B. With a projected separation of 4.4~$\pm$~0.6~AU and an estimated orbital period of 21~$\pm$~5~yr, GJ 3629~AB is likely to yield a dynamical mass in the next several years, making it one of only a handful of brown dwarfs to have a measured mass and an age constrained from the stellar primary.
|
{\label{sec:intro}} Following a decade of attempts to directly image extrasolar giant planets with increasingly sensitive instruments and speckle suppression techniques (e.g., \citealt{Lowrance:2005p18287}; \citealt{Biller:2007p19401}; \citealt{Lafreniere:2007p17991}; \citealt{Liu:2010p21647}), several planetary systems have finally been found through high contrast imaging over the past few years. These discoveries have opened up an exciting new era of planetary science where the atmospheres of non-transiting extrasolar planets can be directly studied for the first time. Five unambigious planets have been imaged orbiting two stars: one around the A5 star $\beta$~Pic (\citealt{Lagrange:2009p14794}; \citealt{Lagrange:2010p21645}), and four surrounding the A5 star HR~8799 (\citealt{Marois:2008p17990}; \citealt{Marois:2010p21591}).\footnote{Because the optically-detected companion to Fomalhaut (\citealt{Kalas:2008p18842}) has not been recovered in the infrared (\citealt{Marengo:2009p18897}), it is unclear whether the object is a planet, perhaps with a large high-albedo ring system, or something else, like a dust cloud from a recent planetesimal collision (\citealt{Janson:2012p23488}). Recently \citet{Kraus:2012p23492} have discovered what appears to be a young ($\sim$2~Myr) accreting giant planet orbiting the transition-disk star LkCa~15 at $\sim$15--20~AU, but a clear understanding of this system, including the mass of the companion, is still lacking. In addition to these objects, a handful of other planetary-mass companions have been found orbiting stars from hundreds (e.g., 1RXS~J1609--2105~b: \citealt{Lafreniere:2008p14057}; GSC~06214--00210~b: \citealt{Ireland:2011p21592}) to thousands (Ross 458~C: \citealt{Goldman:2010p22044}, \citealt{Scholz:2010p20993}; WD~0806-661~B: \citealt{Luhman:2011p22766}) of AU, but the formation mechanism of these enigmatic companions remains obscure (e.g., \citealt{Bowler:2011p23014}).} The fact that all these planets orbit high-mass stars might at first suggest that giant planet formation is more efficient around massive stars, which is a well-established trend observed at smaller separations from radial velocity planet searches (\citealt{Johnson:2007p169}; \citealt{Lovis:2007p17712}; \citealt{Bowler:2010p19983}; \citealt{Johnson:2010p20950}). However, high contrast imaging searches have mostly neglected low-mass stars. Until recently there has been a scarcity of known nearby young M dwarfs, making it difficult to produce statistical comparisons of planet occurrence rates as a function of stellar mass. As a result, even though M dwarfs outnumber AFGK stars by a factor of 2--3 in the solar neighborhood (\citealt{Henry:1997p22864}; \citealt{Bochanski:2010p23010}), our understanding of planet formation is the weakest in this stellar mass regime. We are conducting a high contrast adaptive optics imaging survey of young M dwarfs with Keck-II and Subaru to search for young giant planets and brown dwarfs and measure the frequency of giant planets orbiting low-mass stars. Our targets are newly identified M dwarfs with ages $<$300~Myr and distances $\lesssim$30~pc which were selected based on elevated levels of X-ray and UV emission (\citealt{Shkolnik:2009p19565}; \citealt{Shkolnik:2011p21923}). Compared to high-mass stars, M dwarfs present several advantages as targets for direct imaging. Their higher space densities mean they are on average closer than more massive stars, so smaller physical separations can be probed. Moreover, because M dwarfs are intrinsically faint, direct imaging can detect lower planet masses in the contrast-limited regime. This allows us to reach typical planet masses of a few \Mjup \ at separations of $\sim$10~AU, making our Planets Around Low-Mass Stars (PALMS) survey one of the deepest direct imaging planet searches to date. The first discovery from our PALMS survey was an L0 substellar companion to the young M~dwarf 1RXS~J235133.3+312720 separated by $\sim$120~AU (\citealt{Bowler:2012p23698}). Here we present the discovery of a 0$\farcs$2 companion to the young M3.0 star GJ~3629. \citet{Shkolnik:2009p19565} identified GJ~3629 as a young star based on its high fractional X-ray luminosity, which is comparable to members of the Pleiades ($\sim$125~Myr) and young moving groups (10--100~Myr). \citet{Shkolnik:2009p19565} obtained a high resolution optical spectrum of GJ~3629~A and did not detect Li, thereby setting a lower limit to its age. Based on these constraints they estimate an age of 25--300~Myr, which places our new companion GJ~3629~B below the hydrogen burning minimum mass at the estimated distance of the system (22~$\pm$~3~pc; Section~\ref{sec:dist}).
|
{\label{sec:discussion}} With an angular separation of $\approx$200~mas, GJ~3629~AB has a projected physical separation of 4.4~$\pm$~0.6~AU at a distance of 22~$\pm$~3~pc. \citet{Dupuy:2011p22603} calculate statistical correction factors to convert projected separations of visual binaries into semimajor axes using a variety of input eccentricity distributions. Assuming their conversion factor of 1.16, which represents the case of no discovery bias and an input eccentricity distribution from the observed population of very low-mass binaries, we estimate the orbital period of the GJ~3629~AB system to be 21~$\pm$~5~yr. This raises the possibility of measuring a dynamical mass for the system on a modest timescale. Radial velocity monitoring of the system will also be feasible because the velocity semi-amplitude induced by the companion is expected to be $\sim$1.0~km~s$^{-1}$ assuming an inclination of 90$^{\circ}$. A parallax measurement and continuing astrometric and radial velocity monitoring should be a high priority for this system. A growing list of brown dwarfs now have dynamical mass measurements (e.g., \citealt{Liu:2008p14548}; \citealt{Dupuy:2010p21117}; \citealt{Konopacky:2010p20823}), but very few of these also have good age and metallicity constraints from being companions to well-characterized stars or members of coeval clusters. This rare class of brown dwarfs--- representing both ``age benchmarks'' and ``mass benchmarks''--- functions as an excellent tool to accurately test substellar evolutionary models (\citealt{Liu:2008p14548}). For example, \citet{Dupuy:2009p15627} measured the dynamical mass of the binary brown dwarf HD~130948~BC and used the precise age determination of the primary star (0.8~$\pm$~0.2~Gyr) to test two of the most commonly used substellar cooling models (\citealt{Burrows:1997p2706}; \citealt{Chabrier:2000p161}). They found that these models underpredict the luminosities of HD~130948~BC by a factor of 2--3. Other brown dwarfs with dynamical masses and well-constrained ``environmental'' ages (that is, not dependent on substellar cooling models) are the young ($\sim$1~Myr) eclipsing binary brown dwarf pair 2MASS~J05352184--0546085~AB (\citealt{Stassun:2006p23491}) and the 2.2~$\pm$~1.5~Gyr system HR~7672~B (\citealt{Liu:2002p18017}; \citealt{Crepp:2011p23186}).\footnote{The other systems consisting of one or more brown dwarfs with a dynamical mass measurement and a stellar primary all have poor age constraints. The triple system GJ~569 Bab (\citealt{Martin:2000p20363}; \citealt{Lane:2001p22845}) appears to have an age $\lesssim$1~Gyr but the estimates for the primary star vary widely in the literature (see \citealt{Dupuy:2010p21117} for a summary). Likewise, various age estimates for the $\epsilon$~Indi~Bab triple system (\citealt{Scholz:2003p22538}; \citealt{Mccaughrean:2004p22578}; \citealt{Cardoso:2009p11371}) place it between $\sim$0.5--7~Gyr (see \citealt{Liu:2010p21195} and \citealt{King:2010p23528}). The triple system GJ~802~AB (\citealt{Pravdo:2005p23764}; \citealt{Lloyd:2006p23493}) also has a dynamical mass measurement (\citealt{Ireland:2008p19807}), but the age constraint based on its kinematics is rather weak at $\sim$3--10~Gyr.} Many of the aforementioned benchmarks used unique methods to extract individual masses for the system components. For example, \citet{Dupuy:2009p15627} relied on the HD~130948~BC companions having nearly equal flux ratios to infer relative masses; \citet{Stassun:2006p23491} used the well-constrained inclination of the eclipsing binary 2MASS~J05352184--0546085~AB to extract individual masses from radial velocity curves; and \citet{Crepp:2011p23186} used radial velocity data of the primary star and visual orbit monitoring of the companion to infer the mass of the HR~7672~B. For GJ~3629~AB to join this rare group of benchmarks, individual masses will have to be measured instead of a total mass for the system, which is what relative orbit monitoring yields. A stationary point source is visible $\sim$30$''$ north of GJ~3629~AB in POSS-I and POSS-II images from the Digitized Sky Surveys, but it is not detected by 2MASS. The source (SDSS~J105120.51+360752.3) was identified by \citet{Schneider:2007p23494} as a quasar based on optical spectroscopy from the Sloan Digital Sky Survey (SDSS). Its optical colors from SDSS are very blue, but with a $z$-band magnitude of 18.8~mag it should be possible to use this as a reference object in the near-infrared for absolute astrometry of the GJ~3629~AB system. Finally, we note that several observations can be made to further refine the age of the GJ~3629~AB system: a parallax measurement would enable placement on the HR diagram and resolved near-infrared spectroscopy of GJ~3629~B would provide age constraints though the use of gravity-sensitive features (e.g., \citealt{Allers:2007p66}). Altogether, GJ~3629~AB represents a promising system for future studies.
| 12
| 6
|
1206.3307
|
1206
|
1206.5521_arXiv.txt
|
The FU Orionis (FUor) or EX Orionis (EXor) phenomenon has attracted increasing attention in recent years and is now accepted as a crucial element in the early evolution of low-mass stars. FUor and EXor eruptions of young stellar objects (YSOs) are caused by strongly enhanced accretion from the surrounding disk. FUors display optical outbursts of $\sim$ 4 mag or more and last for several decades, whereas EXors show smaller outbursts ($\Delta$m $\sim$ 2 - 3 mag) that last from a few months to a few years and may occur repeatedly. Therefore, FUor/EXor eruptions represent a rare but very important phenomenon in early stellar evolution, during which a young low-mass YSO brightens by up to several optical magnitudes. Hence, long-term observations of this class of eruptive variable are important to design theoretical models of low-mass star formation. In this paper, we present recent results from our long-term monitoring observations of three rare types of eruptive young variables with the 2-m Himalayan {\it Chandra} Telescope (HCT) and the 2-m IUCAA Girawali Observatory (IGO) telescope.
|
There is now convincing evidence that EXor and FUor outburst phenomena are closely related to the earliest stages of stellar evolution (Herbig, Petrov \& Duemmler 2003). Rarity and obscuration have resulted in a poor understanding of the eruption mechanism in spite of an established integral link to disk accretion (Reipurth \& Aspin 2004; Hartmann, Hinkle \& Calvet 2004). FUor and EXor, also referred to as sub-FUors, are among the most interesting and intriguing class of known pre-main-sequence (PMS) stars. They are likely to be near solar-mass protostars that are still accreting material from their circumstellar disks and are associated with collimated outflows (Sandell \& Weintraub 2001). Only a handful of these objects are known to date (e.g., Vittone \& Errico 2005). Hence, there is an urgent need for observations and studies to facilitate a deeper understanding of their nature and the effects of their associated eruptions. The morphology and nature of small compact reflection nebulae in the star-forming clouds possibly hint at these objects being in a transition phase between an embedded PMS star and a visible Herbig-Haro object (Reipurth \& Bally 2001; Reipurth \& Aspin 2004). A similar case was the emergence of McNeil's nebula, which was found to harbour a possible EXor event (Ojha et al. 2006 and references therein). In this paper, we present optical observations of the post-outburst phases of McNeil's nebula (V1647 Ori), a new reflection nebula in LDN 1415, and a cometary reflection nebula associated with the infrared source IRAS 06068-0641. In Section 2, we present details of the observations and data reduction procedures and in Section 3 we present the results and discuss the short- and long-term variability of three eruptive low-mass young variables.
|
\subsection{McNeil's nebula (V1647 Ori)} The compact source at the base of a variable nebula (McNeil's Nebula Object) in the Lynds 1630 dark cloud in Orion went into outburst in late 2003 (McNeil 2004). Later, the McNeil's Nebula Object was identified as V1647 Ori (Samus 2004). V1647 Ori, a low-mass, deeply embedded, PMS star, has undergone two optical/near-infrared outbursts in the last decade, both of which gradually faded over several months to years. These eruptions are thought to have been the result of large-scale accretion events. Ojha et al. (2006) presented a detailed study of the post-outburst phase of McNeil's nebula using optical ($B,V,R,I$) and near-infrared ($J,H,K$) photometric and low-resolution optical spectroscopic observations. The long-term optical and near-infrared observations showed a general decline in the brightness of the exciting source of McNeil's nebula, V1647 Ori. Our optical images taken in November 2005 showed that V1647 Ori had faded by more than 3 magnitudes since February 2004 (see Figure 2, {\it top panel}). McNeil's nebula itself had also faded considerably. The optical spectra showed strong H$\alpha$ emission with blue-shifted absorption and the Ca II IR triplet (\mbox{8498 \AA}, 8542 \AA~and 8662 \AA) in emission. The presence of the Ca II IR triplet in emission confirmed that V1647 Ori is a PMS star. Therefore, our long-term, post-outburst photometric and spectroscopic observations of V1647 Ori indicated an EXor rather than an FUor event. \begin{figure} \centerline{\includegraphics[width=11cm]{fig2.ps}} \caption{The optical light curve of V1647 Ori in $R$ band. The filled circles show HFOSC and IFOSC measurements from February 2004 to February 2011.} \end{figure} In 2008, V1647 Ori again underwent a strong outburst (Aspin 2008), which is rarely seen in the early phases of low-mass young stellar objects. We monitored V1647 Ori with the HCT and IGO telescopes beginning with the second outburst. No significant variation in brightness of V1647 Ori was seen for about two months since the second outburst began (Ojha et al. 2008). In comparison with the last reported quiescent phase (Ojha et al. 2006), however, there was a brightening of about $\sim$ 3 magnitudes in $R$, and the infrared colors suggested that circumstellar matter of A$_V$ $\sim$ 7.5 mag had probably been cleared during this outburst. Figure 2 ({\it bottom panel}) shows the optical light curve (September 2008 - February 2011) of V1647 Ori in $R$-band. Comparison of the spatial flux distribution of the nebula with the first post-outburst phase in 2004 revealed a change in the dust distribution around the source during the second outburst. From our 2+ year long monitoring observations (September 2008 - February 2011), we see significant short-term variations in the brightness of V1647 Ori since the second outburst began. The source, however, has not faded away considerably as seen in 2004 - 2005. The source magnitude and \mbox{1-$\sigma$} fluctuations in $V$, $R$ \& $I$ during the period of our observations were 18.86$\pm$0.23, 17.04$\pm$0.16, 14.99$\pm$0.12, respectively. The observed properties of the outburst of V1647 Ori are different in several respects from both the EXor and FUor type outbursts, and suggest that this star represents a new type of eruptive young star, one that is younger and more deeply embedded than EXor, and exhibits variations on shorter time scales than FUors. \subsection{LDN 1415 (IRAS 04376+5413)} The new reflection nebula in the not so well studied Lynds opacity class 3 dark cloud LDN 1415 (Lynds 1962) was first detected by Stecklum (2006) in early April 2006 in the vicinity of IRAS 04376+5413. Stecklum, Melnikov \& Meusinge (2007) later reported the presence of a new compact arc-shaped nebula with a size of 20 arcsec in the CCD images of the dark cloud LDN 1415. The brightness peak of the nebula is within the positional error ellipse of IRAS 04376+5413. Optical spectra of the nebula taken by Stecklum, Melnikov \& Meusinger on 21 September 2006 revealed the presence of a P-Cygni profile in the H$\alpha$ line, indicating clear evidence for an FUor or EXor-type outburst due to temporarily enhanced accretion. Kastner et al. (2006) observed this eruptive source with the {\it Chandra} X-ray Observatory's Advanced CCD Imaging Spectrometer imaging array (ACIS-I). No X-ray sources were detected, which constrained the X-ray luminosity of the emergent source to be less than $\sim 2 \times 10^{28}$ erg $s^{-1}$, assuming the distance to the LDN 1415 cloud to be 170 pc. To study the post-outburst phase of the embedded source in the LDN 1415 nebula, we have been carrying out optical observations of this source with HCT and IGO telescopes. We present in Figure 3 variability measurements of the LDN 1415 nebula for a duration of about two and half years. In comparison with the available pre-outburst photometry from POSS II (epoch December 1996) and the KISO (January 2001) quoted in Stecklum, Melnikov \& Meusinger (2007), our first post-outburst data point shows an enhancement of $\sim$ 3.4 mags in the $I$-band. Following this observation, a general decline in the brightness is seen in all three ($VRI$) optical light curves (see Figure 3). Superimposed on this decline, we see the presence of small-scale fluctuations of $\sim$ 0.2 - 0.3 mags over short time scales of 3 - 8 months. This variation is consistent with the young and eruptive nature of this class of objects. Therefore, our long-term, post-outburst optical and NIR photometric and optical spectroscopic monitoring of the LDN 1415 nebula and its associated outburst source from October 2006 to March 2009 (Pawade et al. 2010) suggest an EXor or FUor event, possibly by the least luminous member of the known sample of FUor and EXor objects (Stecklum, Melnikov \& Meusinger 2007). \begin{figure} \centerline{\includegraphics[width=11cm]{fig3.ps}} \vspace*{-5.5cm} \caption{The optical light curve of L1415 nebula in $R$ band. The filled circles show our HFOSC and IFOSC measurements (October 2006 - March 2009). The empty circles show the photometric measurements from Stecklum, Melnikov \& Meusinger (2007).} \end{figure} \subsection{IRAS 06068-0641} A possible FUor-type eruption from the infrared source IRAS 06068-0641 was discovered by the Catalina Real-time Transient Survey (CRTS) on 10 November 2009 (Wils et al. 2009). The object lies inside a dark nebula to the south of the Monoceros R2 association, and is likely related to it. Figure 4 shows the optical light curve (November 2009 - March 2011) of IRAS 06068-0641 in $R$-band. After a significant increase in the brightness from at least late November 2009 ($R$ $\sim$ 13.4 mag) to January 2010 ($R$ $\sim$ 12.2 mag), a general decline in the source brightness can be seen in recent observations. The FUor class of eruptive low-mass YSOs display outbursts of $\sim$ 4 mag or more that last for several decades. EXors show smaller outbursts ($\Delta$m $\sim$ 2 - 3 mag) that last from a few months to a few years and may occur repeatedly (Herbig 1977; Bell et al. 1995; Hartmann 1998). For the more than a year that we followed the source IRAS 06068-0641 (January 2010 - March 2011), the brightness decreased by more than 3.5 mag in $R$ band and it is probably returning to its pre-outburst state. It is therefore possible that we witnessed EXor behaviour, since the source variability had about the correct amplitude (Reipurth \& Aspin 2004; Ojha et al. 2005, 2006). Further photometric observations of this object are required to understand and classify the outburst happening now in IRAS 06068-0641. \begin{figure} \centerline{\includegraphics[width=11cm]{fig4.ps}} \vspace*{-5.5cm} \caption{The optical light curve of IRAS 06068-0641 in $R$ band.} \end{figure}
| 12
| 6
|
1206.5521
|
1206
|
1206.0903_arXiv.txt
|
In this paper the effects of time-dependent Newton constant $G$ during inflation are studied. We present the formalism of curvature perturbations in an inflationary system with a time-dependent Newton constant. As an example we consider a toy model in which $G$ undergoes a sudden change during inflation. By imposing the appropriate matching conditions the imprints of this sharp change in $G$ on curvature perturbation power spectrum are studied. We show that if $G$ increases (decreases) during the transition the amplitude of curvature perturbations on large scales decreases (increases). In our model with a sudden change in $G$ a continuous sinusoidal modulations on curvature power spectrum is induced. However, in a realistic scenario in which the change in $G$ has some finite time scale we expect these sinusoidal modulations to be damped on short scales. The generated features may be used to explain the observed glitches on CMB power spectrum. This puts a bound on $\Delta G$ during inflation of roughly the same order as current bounds on $\Delta G$ during the entire observed age of the universe.
|
There are strong evidences that Newton gravitational ``constant'' $G$ did not change considerably during most of the history of our universe. The evidences range from studies of orbits of planets to pulsar timing, evolution of stellar objects, the cosmic microwave background (CMB) and the effect on the abundance of light elements formed during the period of big bang nucleosynthesis. These studies yield a bound $|\dot{G}/G|\lesssim10^{-(11\hbox{--}13)}\,{\rm yr}^{-1}$ \cite{Uzan:2002vq,Uzan:2010pm}. This seems to be a strong bound but actually it corresponds to $|\Delta G/G| \lesssim 0.1$ over the entire history of our universe, which makes it less dramatic. In addition, it should be noted that many other assumptions are hidden in most of these studies (like taking the other `constants' to be constant). Of course, it is always possible to come up with a complicated theory of variable $G$ that explains all these observations, but least deviations from the standard working model is usually preferred. Thus the continued lack of evidence for a variable $G$ has been convincing for most people to accept $G=$constant. There is, however, no direct indication from any of the mentioned experiments that $G$ has to be constant at very early times, for example, during inflation. This is one of our motivations in this paper to pursue implications of the change in Newton constant during the inflationary era. Inflation has emerged as the leading paradigm for early universe and structure formation \cite{Guth:1980zm} which is strongly supported by cosmological observations \cite{Komatsu:2010fb}. The simplest models of inflation predict almost scale invariant, almost Gaussian and almost adiabatic perturbations on the CMB. In this work we would like to examine the effects of a time-dependent Newton constant in inflationary predictions. We present the formalism for curvature perturbations in inflationary models with a time-dependent $G$. As a particular example, we consider the toy model in which the reduced Planck mass $\Omega \equiv 1/8\pi G$ undergoes a sharp change from $\Omega= \Omega_-$ to $\Omega= \Omega_+$ in which both $\Omega_\pm$ are constant. We put this change in $\Omega$ at an early stage of inflation so the effects of change in Planck mass is within the CMB observational window. As a result one expects to find local features to be imprinted on curvature perturbations from this sudden change in Planck mass. Indeed there has been a lot of interest in the literature to consider the effects of local features in inflation. These features may originate from a sudden change in slow-roll conditions, sudden change in the inflaton mass, particle creation during inflation, field annihilations during inflation, change in sound speed of perturbations or change in fluids equation of state \cite{Starobinsky:1992ts, Leach:2001zf, Adams:2001vc, Gong:2005jr, Joy:2007na, Joy:2008qd, Chen:2006xjb, Chen:2008wn, Chen:2011zf, Chen:2011tu, Hotchkiss:2009pj, Arroja:2011yu, Adshead:2011jq, Abolhasani:2012px, Arroja:2012ae, Romano:2008rr, Battefeld:2010rf, Firouzjahi:2010ga, Battefeld:2010vr, Barnaby:2009dd, Barnaby:2010ke, Zarei:2008nr, Biswas:2010si, Silverstein:2008sg, Flauger:2009ab, Flauger:2010ja, Bean:2008na}. The observational motivations behind these models are to address the glitches in curvature power spectrum on scales associated with $\ell \sim 20-40$. The simplest implementation of a variable $G$ is a scalar-tensor theory of gravity. In fact, there are many theoretical models, like string theory, supergravity and theories of extra dimensions, which involve scalar fields non-minimally coupled to a rank-2 tensor. Thus there is a natural place for theories of variable $G$ within the framework of fundamental physics. But to maintain a constant $G$, and hence respect the null experiments mentioned above, authors often choose to change variables to another rank-2 tensor which {\emph is} minimally coupled to matter. In other words, the metric that describes the geometry of space-time is taken to be the one that appears in the Einstein frame. We will not do so. Instead we work with a model where the effective gravitational coupling, given by $\Omega_+$, has no variation during the standard hot big bang cosmology after inflation but experiences a change during the inflationary period. Having this said, if one insists on performing the field redefinition and goes to the Einstein frame with a constant $G$ throughout, then the nature of the physical phenomena is the same in both frames; it is only a matter of interpretation of what one measures. In particular, the observers may use different clocks and rulers to measure physical processes in these two frames, though the final conclusion is the same for both observers. The effects of time-dependent Newton constant during inflation via a conformal coupling with multiple inflation fields to gravity was recently studied in \cite{White:2012ya}, see also \cite{Deruelle:2010ht, Makino:1991sg, Tsujikawa:2004my, Koh:2005qp, Koh:2010kg, Gong:2011qe, Kubota:2011re, Kallosh:2010ug, Linde:2011nh, Ferrara:2010yw, Ferrara:2010in, DeFelice:2011jm, Jamil:2009sq}. The rest of this paper is organized as follows. In Section \ref{background} we present our setup and the background inflation equations. In Section \ref{perturbations} we present the cosmological perturbation theory for general models with a time-dependent Planck mass. In Section \ref{sudden-change} we consider the special case in which $\Omega$ undergoes a sudden jump. By imposing the appropriate matching conditions we find the curvature perturbations power spectrum which will be used to put some rough bound on $\Delta G/G$ during inflation in this toy model. Discussions are given in Section \ref{discussions} followed by some technical issues in Appendix A.
|
\label{discussions} In this paper we considered the effects of change in Newton constant $G$ or the reduced Planck mass $\Omega = 1/8 \pi G$ during inflation. The motivations for considering a time-dependent Planck mass may come from models of high energy physics, like string theory, in which the effective four-dimensional Planck mass is a function of extra dimension fields such as the volume modulus. In the standard big bang cosmology there are tight constraints on the variations of $G$. However, there is no direct constraint on the variation of $G$ during inflation. In this work we presented the formalism of cosmological perturbation theory in the models with a time-dependent $\Omega$. In particular, we have seen that a time-varying $\Omega$ causes gravitational anisotropy in which the Bardeen potential and the Newton potential are not equal. This can have interesting implications for the CMB and for lensing. Also, in the context of the separate universe approach, we have seen that a time-dependent $\Omega$ effectively acts as a new source of entropy perturbations. As a result, $\zeta$ does change on super-horizon scales even for an adiabatic inflationary fluid. As a specific solvable example, we considered a toy model in which $\Omega$ undergoes a rapid change from $\Omega_-$ to $\Omega_+$ during inflation. After imposing the appropriate matching conditions on comoving curvature perturbations ${\cal R}_c$ and its derivative $\dot {\cal R}_c$ we obtained the outgoing power spectrum in terms of the incoming power spectrum. The effects of change in $\Omega$ is encoded in the transfer function $T(k)$. On physical grounds we expect that $T(k)$ should have localized features centered around $k=k_c$, where $k_c$ represents the mode which leaves the horizon at the time of phase transition. This is verified in our analysis. In addition, we also see a continuous sinusoidal modulation on $T(k)$. We argued that this is an artifact of our simplifying assumptions that the change in $\Omega$ takes place instantaneously. In realistic models in which the change in $\Omega$ takes some finite non-zero time one expects that for very small scales $T(k)$ reaches the asymptotic value unity. On the other hand, if one chooses to work in the Einstein frame with a constant $G$, then it is natural to ask what mechanisms create these features on power spectrum. As we discussed in Introduction and in Sec.~\ref{background}, in the Einstein frame the matter sector will have a complicated non-standard Lagrangian. In this view, in the Einstein frame one expects to see some sudden changes in fluid properties which causes the local features on power spectrum. In some ways, these sudden changes in fluid or matter sectors may be modeled via any of the mechanisms studied so far in literature, i.e., in \cite{Starobinsky:1992ts, Leach:2001zf, Adams:2001vc, Gong:2005jr, Joy:2007na, Joy:2008qd, Chen:2006xjb, Chen:2008wn, Chen:2011zf, Chen:2011tu, Hotchkiss:2009pj, Arroja:2011yu, Adshead:2011jq, Abolhasani:2012px, Arroja:2012ae, Romano:2008rr, Battefeld:2010rf, Firouzjahi:2010ga, Battefeld:2010vr, Barnaby:2009dd, Barnaby:2010ke, Zarei:2008nr, Biswas:2010si, Silverstein:2008sg, Flauger:2009ab, Flauger:2010ja, Bean:2008na} in the context of Einstein frame gravity. In our analysis so far, we have not presented a dynamical mechanism which causes the variations in $\Omega$. One interesting realization for creating a dynamical time-dependence in $\Omega$ is to couple the gravitational system non-minimally to scalar fields which vary during inflation. In this view, $\Omega$ can couple either to the inflaton field or to additional light or heavy fields present in the model. As a concrete example, consider an inflationary model with two fields: the inflaton field $\phi$ and the waterfall field $\chi$ with the matter Lagrangian \begin{equation} {\cal L}_M = -\frac{1}{2} (\partial\chi)^2 -\frac{1}{2} (\partial\phi)^2 - V(\chi, \phi). \end{equation} Here the potential $V$ is the same as in models of hybrid inflation \cite{Linde:1993cn, Copeland:1994vg} \begin{equation} V = \frac{1}{2}m^2\phi^2 + \frac{\lambda}{4\lambda}(\chi^2 - \frac{M^2}{\lambda})^2 + \frac{1}{2}g^2\phi^2\chi^2 \, . \end{equation} Furthermore, suppose that \ba \Omega = \Omega_- + \xi \chi^2 \, , \ea in which $\xi$ is a dimensionless number. The picture we have for the dynamics of the system is similar to waterfall dynamics in hybrid inflation. The system has two distinguished periods, the time before the waterfall phase transition corresponding to $\phi < \phi_c= M/g$ and and the time after the waterfall, $\phi > \phi_c$. The waterfall field is very heavy during inflation so it is frozen during the first stage of inflation and $\chi=0$. As a result, during the first stage of inflation $\Omega = \Omega_-$. After $\phi$ reaches the critical value $\phi= \phi_c$, the waterfall field becomes tachyonic and rapidly rolls to its global minima $\chi= \pm M/\sqrt \lambda $. The quantum fluctuations of the waterfall field $\delta \chi$ play crucial roles as studied in recent works \cite{Lyth:2010ch, Abolhasani:2010kr, Fonseca:2010nk, Abolhasani:2011yp, Gong:2010zf, Lyth:2012yp, Abolhasani:2012px, Clesse:2010iz, Martin:2011ib, Mulryne:2011ni, Avgoustidis:2011em, Kodama:2011vs, Abolhasani:2010kn, Bugaev:2011qt, Bugaev:2011wy}. As a result, at the end of waterfall transition, on each Hubble patch $\chi^2 = \langle \delta \chi^2 \rangle = M^2 /\lambda $ and \ba \Omega_+ = \Omega_- + \frac{\xi M^2}{\lambda} \, . \ea One can make the waterfall phase transition sharp enough, so the time of change in $\Omega$ is reasonably short, say one e-folding or so. However, it is crucial to note that it is not arbitrarily sharp and in principle it takes a finite non-zero time scale for the waterfall phase transition to complete. This is expected to eliminate the unwanted non-decaying sin modulations in $T(k)$ for small scales. Also note that, depending on model parameters \cite{Abolhasani:2011yp, Abolhasani:2012px}, the waterfall transition can happen either during early stages of inflation or towards the end of inflation. Having presented this dynamical mechanism for change in $\Omega$, it is an interesting exercise to study this model in details and see how the waterfall dynamics can be employed to induce a time-dependence in Planck mass. We would like to come back to this question in a future work.
| 12
| 6
|
1206.0903
|
1206
|
1206.0597_arXiv.txt
|
In this paper we report accurate \textit{Chandra} positions for two ultraluminous X-ray sources: NGC 7319-X4 at Right Ascension (RA) = 339.02917(2)$^{\circ}$, Declination (Dec) = 33.97476(2)$^{\circ}$ and NGC 5474-X1 at RA = 211.24859(3)$^{\circ}$, Dec = 53.63584(3)$^{\circ}$. We perform bore-sight corrections on the \textit{Chandra} X-ray Satellite observations of these sources to get to these accurate positions of the X-ray sources and match these positions with archival optical data from the Wide Field and Planetary Camera 2 on board the Hubble Space Telescope. We do not find the optical counterparts: the limiting absolute magnitudes of the observations in the WFPC2 standard magnitude system are B = $-7.9$, V = $-8.7$ and I = $-9.3$ for NGC 7319-X4 and U = $-6.4$ for NGC 5474-X1. We report on the X-ray spectral properties and we find evidence for X-ray variability in NGC 5474-X1. Finally, we briefly discuss several options for the nature of these ULXs.
|
Ultraluminous X-ray sources (ULXs) are pointlike, off-nuclear X-ray sources that have a luminosity that is larger than the Eddington luminosity of a 10 M$_{\sun}$ black hole ($L_{\textrm{Edd}}(10 \textrm{ M}_{\sun}) \approx 10^{39}$ erg s$^{-1}$; \citealt{colbert05}). Several scenarios have been proposed to explain the high luminosities of these sources. The sources with $L \lesssim 10^{40}$ erg s$^{-1}$ can be explained by stellar mass black holes that emit radiation beamed instead of isotropically \citep{king01}. Alternatively, the black hole can be more massive than the typical 10 M$_{\sun}$ if formed by a low-metallicity progenitor, which would lead to a higher Eddington luminosity \citep{belczynski10}. ULXs with luminosities above $10^{41}$ erg s$^{-1}$, sometimes called hyperluminous X-ray sources, may however need black holes with $M \gtrsim 1000 M_{\sun}$ to explain \citep{colbert05}. Such black holes are referred to as intermediate mass black holes (IMBHs). Another possibility is that some of these sources could be recoiling supermassive black holes (SMBHs). In the standard $\Lambda$CDM cosmology galaxies grow through hierarchical mergers. The central massive black holes of merging galaxies may eventually merge as well. Simulations show that when two black holes coalesce, gravitational waves are emitted that carry away linear momentum in a preferential direction just prior to the merger. As a result the black hole receives a kick in the opposite direction. The size of the kick depends on the magnitude and orientation of the spins of the black holes before the merger and on their mass ratio (\citealt{damour06}, \citealt{ferrarese02}, \citealt{merritt09}). The maximum kick velocity is of the order of 1000 km/s \citep{baker08}, which, with time, would allow one to observe these recoiled SMBHs as bright off-nuclear X-ray sources \citep{jonker10}. Another possibility is that some of these ULXs are supernovae. SN type IIn are the X-ray-brightest supernovae and therefore the best candidates \citep{immler03}. Finally, some ULXs could potentially be explained as the remnant BH from the smaller galaxy that underwent a merger event with a larger galaxy. Finding the optical counterparts of ULXs can help us distinguish between the different scenarios for their nature. If the ULX is a stellar or intermediate mass black hole we expect to see an accretion disc and/or a companion star. A recoiling SMBH would be surrounded by a cluster of stars \citep{merritt09}, whereas a supernova could be distinguished by its decaying light curve. For several ULXs the optical counterpart has been reported (e.g. \citealt{motch11}, \citealt{yang11}, \citealt{wiersema10}). In this Paper, we investigate two ULXs that were observed by the \textit{Chandra} X-ray Observatory and the Hubble Space Telescope (HST). NGC 7319-X4 \citep{liu11} is located in the late-type galaxy NGC 7319, one of the galaxies in Stephan's Quintet. NGC 5474-X1 \citep{swartz11} is the brightest X-ray source in the highly asymmetrical late-type galaxy NGC 5474.
|
Using bore-sight corrected \textit{Chandra} images we determine accurate positions for the ULXs NGC 7319-X4 and NGC 5474-X1. We also improve the astrometry of HST observations of these galaxies and search for optical sources at the ULX positions. The position of NGC 7319-X4 is RA = 339.02917$^{\circ}$, Dec = 33.97476$^{\circ}$, with a 1-$\sigma$ error circle of 0.3'' on the HST image. We find the nearest star at 0.8'' from this position -- at the distance of NGC 7319 (90.75 Mpc) this corresponds to $\sim$ 350 pc. The position of NGC 5474-X1 is RA = 211.24859$^{\circ}$, Dec = 53.63584$^{\circ}$, with a 1-$\sigma$ error circle on the HST image of 0.22''. Again the nearest star that we find lies at 0.8'' from this position. This corresponds to $\sim$ 25 pc at the distance of NGC 5474. We do not find the optical counterparts to the ULXs in NGC 5474 and NGC 7319 in the HST data, but using our limiting absolute magnitudes we can put constraints on their counterparts. To constrain the spectral type of a possible companion star we use table 15.7 from Allen's astrophysical quantities \citep{cox_book}. This table contains absolute V-magnitude and color calibrations for the MK spectral classes. We use the values from the WFPC2 Cookbook\footnote{www.stsci.edu/hst/wfpc2/analysis/wfpc2\_cookbook.html} to convert these UBVRI magnitudes to WFPC2 standard magnitudes. \subsubsection{NGC 7319} The limiting magnitudes for the source in NGC 7319 are $-7.9$ in the WFPC2 B-band, $-8.7$ in the WFPC2 V-band and $-9.3$ in the WFPC2 I-band. The brightest stars in the WFPC2 B- and V-band are O9 supergiants with a B-band magnitude of $-7.5$ and a V-band magnitude of $-6.5$, and in the WFPC2 I-band the brightest stars are M5 supergiants with a magnitude of $-8.7$. None of these would have been detected in the HST images. With our limiting absolute magnitudes we can not exclude any type of single star as counterpart. If the ULX is an accreting compact object it will probably have an accretion disc. These can also be visible in optical light -- in fact, disc emission may dominate the emission in the blue part of the spectrum. Can our limiting absolute magnitudes constrain the size of such a disc? \citet{vanparadijs94} found a relation between $\Sigma$ and the absolute V-band magnitude of low-mass X-ray binaries (LMXBs). $\Sigma$ depends on the Eddington ratio and period of the system as $\Sigma = (L_X/L_{Edd})^{1/2} (P/1 hr)^{2/3}$. Using this relation and the limiting V-band magnitude of $-8.7$ yields an upper limit for $\log(\Sigma)$ of 4.5 (with the caveat that we are extrapolating this relation from $\log(\Sigma) = 2.01$, the maximum value in the sample of \citet{vanparadijs94}). This is not very constraining -- if we assume that the system is accreting at the Eddington limit we find a maximum period of 640 years, much longer than that of any LMXB. This does not give us a useful constraint on the size of a possible accretion disc. At these high X-ray luminosities X-ray irradiation of the donor star and accretion disc is likely to play an important role (\citealt{copperwheat05}, \citealt{copperwheat07}, \citealt{patruno08}). \citet{copperwheat05} give absolute V-magnitudes for an irradiated O5V star and G0I star with irradiated disc around a 10, 100 and 1000 M$_{\odot}$ black hole (their figure 6). The ULX in NGC 7319 has a hardness ratio of $\sim$ 1.4. An absolute limiting V-band magnitude of $-8.7$ means that an O5V star would not have been detected regardless of the mass of the black hole, but a G0 supergiant around a 100 or 1000 M$_{\odot}$ black hole would have been visible. The most X-ray bright supernovae, type IIn, are also bright at optical wavelengths. The average intrinsic B-band magnitude at maximum brightness of these SN is $M_B = -19.15 \pm 0.92$ \citep{richardson02}, and they fade on timescales of years \citep{filippenko97}. Figure 4 of \citet{tsvetkov08} compares the lightcurves of 5 SN IIn, showing that M$_V \leq -15$ for more than 400 days. The X-ray luminosity peaks around 400-1000 day after the explosion. This makes the supernova scenario unlikely for this source. The HST observations were made a year before the first X-ray observation. If the supernova had gone off before the HST observation, it would have been detected in the B-band images. So the only possibility is that it went off after the HST observations, but in time for it to be bright in X-rays a year later, leaving a very short period of time in which the explosion could have occurred. Since NGC 7319 is a Seyfert 2 galaxy and hence has a detected nuclear X-ray source, the ULX is probably not a recoiling supermassive black hole. Another scenario that we have to investigate is that the X-ray source may be a background quasar. If the source is situated behind NGC 7319, we expect a significant contribution of that galaxy to the column density of neutral hydrogen on top of the Galactic value. We assume a typical disc height of 100 parsec and an average density of 1 particle per cm$^3$ \citep{ferriere01}. The contribution of a face-on spiral galaxy to the column density is then $\sim 3 \times 10^{20}$ cm$^{-2}$, which gives a visual extinction $A_V \approx 0.2$ (using the relation of \citealt{predehl95}). The foreground Galactic extinction in the direction of NGC 7319 is $A_V = 0.26$; adding the extra extinction from NGC 7319 gives $A_V \approx 0.46$. This is lower than, although consistent with, the value that we find from the X-ray spectrum, $A_V = 0.9 \pm 0.4$. Variability in the X-ray spectrum can also help to exclude the AGN scenario. We do not find significant variability between the two \textit{Chandra} observations; therefore we cannot refute the possibility that this source is a background AGN based on X-ray variability. Another clue about the background nature of the source can be gained from the X-ray to optical flux ratio. Most \textit{Chandra}-selected AGN have an X-ray (0.5-2.0 keV) to optical (R-band) flux ratio smaller than 10 (\citealt{barger03}, \citealt{laird09}). For this source however, this ratio is larger than 70 in the V-band, larger than 74 in the B-band and larger than 90 in the I-band. The optical flux limits have been corrected for absorption. These values make the background AGN scenario less likely for this source. \subsubsection{NGC 5474} For NGC 5474 we have a limiting absolute magnitude of $-6.4$ in the WFPC2 U-band. This means that the counterpart can not be an O type main sequence star, nor an A-type or earlier supergiant. Those stars have a WFPC2 U-band magnitude lower than $-6.4$ so they should have been detected in the HST image. Assuming that $F_{\lambda} \propto \lambda^{-2}$ for accretion discs gives M$_U$ - M$_V \approx -1$, yielding a limiting V-band magnitude of approximately $-5.4$ for this source. Using the relation of \citet{vanparadijs94}, this limiting magnitude infers a maximum value of $\log\Sigma$ of 3.1, or, assuming that the ULX is accreting at its Eddington limit, a maximum period of $\sim$ 5 years (again, we are extrapolating the relation between $\Sigma$ and M$_V$). This is such a high upper limit that an LMXB accretion disc would not be detected in these images. X-ray irradiation of the disc and donor star may increase the optical luminosity of the system. The combination of an increasing X-ray luminosity and the non-detection of an optical counterpart in the U-band three years after the first \textit{Chandra} observation makes it very unlikely that NGC 5474-X4 is a supernova. To make that scenario work the first \textit{Chandra} observation (in September 2006) would have had to catch the very steep rise of the X-ray luminosity about one year after the supernova \citep{immler03}, but in that case it would have been visible in the U-band HST image three years later \citet{tsvetkov08}. On top of that, a SN IIn in NGC 5474 would have a peak apparent magnitude m$_V \approx 10$. It seems very unlikely that such a source went unnoticed. Since we do not observe a nuclear X-ray source in NGC 5474 we cannot exclude the recoiling SMBH scenario on that ground. In the case of a recoiling SMBH the optical counterpart would probably be related to the material that remains gravitationally bound to the black hole -- the accretion disc, broad line region and part of the nuclear star cluster -- and its luminosity would be comparable to that of an AGN. The absolute magnitudes of AGN vary, but the faintest that have been reported thus far have absolute magnitudes M$_B$ = -9.8 for NGC 4395 \citep{filippenko89} and -11.6 for NGC 3031/M81 \citep{ho97}. With a limiting U-band magnitude of $-6.4$ NGC 5474-X1 is much fainter than these faintest AGNs, making it unlikely that it is a recoiling SMBH. The Galactic absorption in the direction of NGC 5474 is only $A_V = 0.036$ mag, while we find an absorption of $A_V = 0.54 \pm 0.08$ mag from the X-ray spectrum. This means there is significant absorption on top of the Galactic foreground, possibly local to the source or because the source is situated behind NGC 5474. The two \textit{Chandra} observations, spaced 15 months apart, show similar spectra but with a factor 5 increase in luminosity towards the later observation. This does not exclude a background AGN. It does mean that this is indeed a single source instead of e.g. a group of LMXBs. However, the X-ray (0.5-2.0 keV) to optical (U-band) flux ratio of this source is larger than 500, while typically AGN have a ratio smaller than 10 in the R-band. This argues strongly against the background AGN scenario. \subsection{Conclusions} We can not exclude any type of single star as the companion of the ULX in NGC 7319. The limiting magnitudes for this source also allow an LMXB with accretion disc. \citet{copperwheat05} calculated V-band magnitudes for X-ray irradiated discs and donor stars. By comparing these with our limiting magnitude we can exclude a G0 supergiant around an IMBH of more than 100 M$_{\odot}$. The source could also be a supernova, although this is only possible if it went off in a short period of time right after the HST observations. The recoiling black hole scenario is not very likely for this source, given that NGC 7319 also has an AGN. The source could be a background AGN but the X-ray to optical flux ratios argue against this scenario. The companion of the ULX in NGC 5474 cannot be a blue supergiant or O-type main sequence star. Other stellar types and an LMXB with accretion disc are still an option. The source is probably not a supernova, since the increasing X-ray luminosity and non-detection in the HST observation make this very unlikely. It can be a recoiling black hole, but then its optical luminosity is much lower than is usual for AGN. There is significant absorption on top of the Galactic foreground absorption in the direction of NGC 5474 which could point to a background object. However, the X-ray to optical flux ratio is much higher than usual for AGN, which argues strongly against this scenario. Finding the optical counterpart or measuring variability of the X-ray source can help decide whether or not this is a background object.
| 12
| 6
|
1206.0597
|
1206
|
1206.2889_arXiv.txt
|
\baselineskip=18pt We propose a new, generic mechanism of inflation mediated by a balance between potential forces and a Chern-Simons interaction. Such quasi-topological interactions are ubiquitous in string theory. In the minisuperspace approximation, their effect on the dynamics can be mapped onto the problem of a charged particle in an electromagnetic field together with an external potential; slow roll arises when the motion is dominated by the analogue of `magnetic drift'. This mechanism is robust against radiative corrections. We suggest a possible experimental signature which, if observed, might be considered strong evidence for string theory.
|
Inflation% ~\cite{Guth:1980zm} is at the same time a spectacular phenomenological success, and an enduring theoretical challenge. It seems clear that the universe underwent a period of rapid expansion in its early history; however, it seems rather difficult to engineer a field-theoretic model which sustains a sufficiently extended inflationary phase. Typical models of inflation involve the evolution of an effective scalar field, the inflaton, along its effective potential. If the motion along the potential is sufficiently slow, so that the energy density is dominated by the inflaton potential energy over all other sources, the local value of the potential acts as an effective cosmological constant, slowly evolving in time. The notion of `sufficiently slow' is quantified by the slow-roll parameters of the Hubble scale $H=\dot a/a$ \bea \label{slowroll} \epsilon_H &=& -\frac{\dot H}{H^2} \nonumber\\ \eta_H &=& -\frac{\ddot H}{2H\dot H} \ . \eea The conditions $\epsilon_H,\eta_H\ll 1$ define the period of slow-roll inflation; one needs this era to last long enough so that the scale factor $a$ undergoes $60$ or more e-foldings of exponential expansion. There are a number of mechanisms that have been proposed to achieve such an extended inflationary era: \begin{description} \item{1.} Make the inflaton potential $\cV$ extraordinarily flat% ~\cite{Albrecht:1982wi,Linde:1981mu}, so that $H^2\sim\cV/3\mpl^2$ changes very slowly over a large region of field space. Unfortunately, scalar field potentials are subject to strong radiative corrections, making such a flat potential generically unnatural. The potential can be protected through a symmetry, for instance making the inflaton a pseudo-Goldstone boson as in `natural' inflation% ~\cite{Freese:1990rb}; however, this approach has its own set of issues -- typically one cannot arrange the periodicity of the axion to be large enough to sustain inflation long enough% ~\cite{Banks:2003sx}.% \footnote{Models employing axion monodromy% ~\cite{McAllister:2008hb} suggest a way around this issue.} \item{2.} Engineer the kinetic term so that the inflaton motion has a limiting velocity and/or small sound speed% ~\cite{ArmendarizPicon:1999rj,Silverstein:2003hf}. Here again, the form of the kinetic energy needed is sensitive to quantum corrections. \item{3.} Arrange for some form of `friction' or damping to impede the evolution of the inflaton% ~\cite{Berera:1995ie}, for instance via particle production. These approaches again involve a degree of fine-tuning, since the produced particles mustn't dominate the energy density while being sufficiently produced to overcome the dilution resulting from the expansion, and at the same time achieving the needed damping; there can also be issues with scale dependence of perturbations. \end{description} In general, there is a consistency issue -- the `eta problem' -- as to why quantum corrections to the effective action of the inflaton $\phi$ don't contribute $\Delta\eta=\Delta m_\phi^2/3H^2\sim O(1)$. Supersymmetry, which can protect scalar masses, doesn't help here -- $H$ is an effective scale of supersymmetry breaking. This article introduces a fourth general mechanism of inflation, which we call `EM-flation'.% \footnote{We would have called it `Chern-Simons inflation' but that name appears to be taken% ~\cite{Alexander:2011hz}.} This term refers to both its intrinsic reliance on the electromagnetic fluxes of antisymmetric tensor (AST) gauge fields, and to a strong analogy with the forced dynamics of a charged particle in a magnetic field. The basic idea is that a Chern-Simons term, linear in time derivatives, can play the same role in field space dynamics that a magnetic field does for a point particle. For the particle in a strong field, other force terms in the equations of motion balance against the magnetic force rather than kinetic terms. Similarly, in the presence of a large Chern-Simons coupling, the resulting field motion is the direct analogue of the `magnetic drift' of a charged particle in a strong magnetic field. One need assume no special features of either the kinetic energy or the potential -- the kinetic term plays only a minor role, and potentials can be steep; all that is required is to be able to make the strength of the Chern-Simons interaction sufficiently large, so that `magnetic drift' is slow, in order to generate a long inflationary era. Because no special property of the effective action is assumed -- other than a large Chern-Simons coupling, whose coefficient is quantized -- the eta problem should be absent. Recently, a model of `chromo-natural' inflation% ~\cite{Adshead:2012kp,Adshead:2012qe} was proposed by two of the authors; as we shall see, this model is the first example of this new inflationary mechanism. Let us summarize the chromo-natural inflationary model. The starting point is natural inflation% ~\cite{Freese:1990rb}, where the inflaton is an axion field $\cX$. The axion is assumed to have a potential generated by some unspecified non-perturbative effects, and interacts with a weakly-coupled $SU(2)$ gauge field via the usual topological density \be \label{chiFF} \lambda\left(\frac{\cX}{\fa} \right) \tr [\cF\wedge \cF] \ . \ee Under the assumption of a texture in the gauge field background, conformal to the spatial dreibein,% \footnote{More generally, the ansatz $\cA_i^{\bar ab} = a\psi J_i^{\bar ab}$, where $J_i^{\bar ab}$ are the generators of SU(2) in the $N$ dimensional representation, allows one to embed the scenario in an arbitrary rank gauge group.} \be \label{Aansatz} \cA_i^a = \psi(t) a(t)\delta_i^a \ , \ee the minisuperspace effective action for the axion $\cX$, scale factor $a$, and gauge field scale $\psi$ is \be \label{CrAct} \cS = \int dt\,a^3 \Bigl[ -m_p^2\Bigl(\frac{\da}{a}\Bigr)^2 + \frac 32(\dot\psi+H\psi)^2 - \frac32 g^2\psi^4 + \half\dot\cX^2 - \mu^4\Bigl(1+\cos\Bigl(\frac{\cX}{\fa}\Bigr)\Bigr) - \frac{3g\lambda\cX}{\fa}\psi^2(\dot\psi+H\psi) \Bigr] ~. \ee The scale factor $a(t)$ undergoes slow roll inflation for sufficiently large axion coupling $\lambda\gg 1$, small gauge coupling $g$, and appropriately tuned scale $\mu$ of the axion potential. The key feature of the equations of motion in this parameter regime is that the axion velocity drives the growth of the gauge field, while the gauge field back-reacts to slow the motion of the axion sufficiently that the slow-roll conditions are satisfied, even when the axion potential isn't particularly flat. In the action, the axion coupling to the gauge field is of the same order as the potential, both of which are much larger than the kinetic energy terms, when evaluated on the inflating trajectory; however, being a topological term, the $\cX \cF{\small\wedge}\cF$ term does not contribute to the stress tensor, and so the equation of state is potential dominated $P\approx-\rho$. The plan for the remainder of this article is as follows: In section 2, we examine the dynamics of a charged particle in a magnetic field as a toy model, showing how the mini-superspace dynamics of chromo-natural inflation embed in this toy model, as well as how its dynamics exhibit the `magnetic drift' phenomenon. In section 3, we consider a sampling of the Chern-Simons terms in string theory, and focus on a particular model involving D7-branes. This is particularly interesting in light of the role that such branes play in recent investigations of string model-building (see% ~\cite{Denef:2008wq} for a review). Section 4 contains a cursory look at how density perturbations are expected to arise in this context. We conclude in section 5 with a summary discussion, comment on the stability and genericity of the mechanism, and propose a potential experimental signature.
|
We have outlined a new mechanism of inflation based on the idea that a Chern-Simons term has the same effect on the effective field space dynamics as a magnetic field does on a charged particle; inflation can then be slow-roll because the inflaton experiences `magnetic drift' forces that balance the force from the inflaton potential, when the effective Chern-Simons coupling is parametrically large. In such a situation, the kinetic term of the effective inflaton plays only a minor role, as the forcing from the potential is compensated largely by the Chern-Simons term. Thus there should be no eta problem; it is not the kinetic term that matters, nor the flatness of the potential, but rather the strength of the Chern-Simons term, which is quasi-topological in nature and therefore rather insensitive to quantum corrections in the effective action. The Chern-Simons coupling is of order one in string units (or Planck units, in M-theory), times an integer; so long as the scale of the inflaton potential is sufficiently small relative to this scale and the compactification scale, a long period of slow-roll inflation results. The examples investigated above bear out this intuition. A general feature of this inflationary scenario is its robustness against quantum corrections to the effective action of the inflaton, which is the source of the eta problem in other approaches. Quantum effects generate corrections to the effective potential and kinetic terms. These corrections spoil any delicate tuning that has been done in order to engineer a long period of inflation, when only these terms in the action are active in the inflationary mechanism. In EM-flation, it doesn't matter whether there are order one corrections of this sort -- the expressions~\pref{etaeps} for the slow roll parameters are indeed order one ratios of the effective potential and its derivatives, times the velocity of the inflaton. However, this velocity is of order $1/\lambda$ due to the influence of `magnetic damping', instead of the value that pertains when only potential forces and Hubble damping are present. By a suitable arrangement of the scales in the problem, and perhaps tuning some discrete parameters, the slow-roll parameters are reliably kept small. How large is the EM-flation basin of attraction? Consider the D7-brane example of section~\ref{D7sect}. In the large $\lambda$ limit, the evolution of $b$, $c$ is much faster than $X$, so we take the latter as frozen at some value during the time that $b$ and $c$ are relaxing to an attractor. The equations of motion for these variables have the form \bea \Gamma(w)\partial_t w_a &=& \partial_{w_a} \cU(w) \nonumber\\ \cU &=&(\hat\lambda\cV') w_1 w_2 + \frac16 \hat\lambda^2 H(w_1^4+w_2^4+7 w_1^2w_2^2)+3\gamma_X H^3(w_1^2+w_2^2) \nonumber\\ \Gamma &=& 9\gamma_X H^2 + \hat\lambda^2(w_1^2+w_2^2) \eea where $w_1=\sqrt\gamma_b\, b$, $w_2=\sqrt{\gamma_c}\,c$, and $\hat \lambda = \lambda/\sqrt{\gamma_a\gamma_b}$. The slow-roll dynamics of $b$, $c$ at fixed $X$ is gradient flow, and a plot of the potential $\cU$ shows that the fixed point at $b=c=0$ is an unstable saddle point, while EM-flation is the only stable attractor of the dynamics. The basic intuition is that, once all the fields are nonzero, the Chern-Simons term is active in the dynamics and the dynamics of the inflaton $X$ drive the AST fields $b$, $c$ to larger values and the EM-flation attractor. So in any example where the EM-flation mechanism is in effect, it applies to {\it all} inflationary trajectories. Similarly, the attractor for chromo-natural inflation is the dominant component of the gauge field configuration space, with only a small region around the origin (whose size scales inversely with $\lambda$) attracted toward $\cA=0$ rather than the chromo-natural attractor. Finally, it seems that a characteristic feature of the EM-flation model of section~\ref{D7sect} is a slight anisotropy of the Hubble expansion rate during inflation, see for example equation~\pref{deltaH}, driven by an expectation value for AST fields. This anisotropy will imprint on the CMB perturbations and lead to a potentially observable signature. In the context of ordinary inflation, the effect of an anisotropic Hubble expansion rate was considered in% ~\cite{Ackerman:2007nb}, with the result that the power spectrum picks up a rotationally non-invariant contribution $\delta\cP/\cP$ of order $(H_3-H_2)/H_{tot}$. This latter quantity was estimated in the example of section~\pref{D7sect} to be generically of order $\epsilon_H/\sqrt{N}$. The fluctuation spectrum anisotropy $\delta\cP/\cP$ might be measured down to the level of one percent by the Planck experiment% ~\cite{Ma:2011ii}.% \footnote{An initial indication of such an effect in the data% ~\cite{Bennett:2010jb} has since been discounted as an instrumental effect% ~\cite{Hanson:2010gu}.} It seems reasonable that the fluctuation spectrum anisotropy of EM-flation is at a level similar to that found in% ~\cite{Ackerman:2007nb} for ordinary inflation, given the similarity of the perturbation calculation of section~\ref{perts}; then a fluctuation anisotropy signal might well be within experimental reach. Needless to say, the detection of such an anisotropy would constitute strong evidence for string theory. \vskip 2cm \noindent{{\bf Acknowledgments:}} Thanks to J. Harvey, W. Hu and J. Marsano for helpful discussions, and to E. Silverstein for comments on the first version of this article. This work was supported in part by DOE grant DE-FG02-90ER-40560, and by the Kavli Institute for Cosmological Physics at the University of Chicago through NSF grants NSF PHY-0114422, NSF PHY-0551142 and an endowment from the Kavli Foundation and its founder Fred Kavli. P.A. and M.W. thank the Aspen Center for Physics for its hospitality and support through National Science Foundation Grant No. 1066293, as this work was nearing completion. \vskip 2cm
| 12
| 6
|
1206.2889
|
1206
|
1206.4307_arXiv.txt
|
Advancements in infrared (IR) interferometry open up the possibility to spatially resolve active galactic nuclei (AGN) on the parsec-scale level and study the circumnuclear dust distribution, commonly referred to as the ``dust torus'', that is held responsible for the type 1/type 2 dichotomy of AGN. We used the mid-IR beam combiner MIDI together with the 8\,m telescopes at the Very Large Telescope Interferometer (VLTI) to observe the nucleus of the Seyfert 2 galaxy NGC~424, achieving an almost complete coverage of the $uv$-plane accessible by the available telescope configurations. We detect extended mid-IR emission with a relatively baseline- and model-independent mid-IR half-light radius of $(2.0\pm0.2)\,\mathrm{pc}\,\times(1.5\pm0.3)\,\mathrm{pc}$ (averaged over the $8-13\,\micron$ wavelength range). The extended mid-IR source shows an increasing size with wavelength. These properties are in agreement with the idea of dust heated in thermal equilibrium with the AGN. The orientation of the major axis in position angle $\sim-27^\circ$ is closely aligned with the system axis as set by optical polarization observations. Torus models typically favor extension along the mid-plane at mid-IR wavelengths instead. Therefore, we conclude that the majority of the pc-scale mid-IR emission ($\ga$60\%) in this type 2 AGN originates from optically-thin dust in the polar region of the AGN, a scenario consistent with the near- to far-IR SED. We suggest that a radiatively-driven dusty wind, possibly launched in a puffed-up region of the inner hot part of the torus, is responsible for the polar dust. In this picture, the torus dominates the near-IR emission up to about $5\,\micron$, while the polar dust is the main contributor to the mid-IR flux. Our results of NGC~424 are consistent with recent observations of the AGN in the Circinus galaxy and resemble large-scale characteristics of other objects. If our results reflect a general property of the AGN population, the current paradigm for interpreting and modeling the IR emission of AGN have to be revised.
|
\setcounter{footnote}{0} The dusty environment around supermassive black holes in active galactic nuclei (AGN) came in reach of direct observations using the new high spatial resolution capabilities of long-baseline infrared (IR) interferometry. Within the last years several observing campaigns revealed emission sources in the nuclear region that shows clear characteristics of AGN-heated dust on the sub-parsec to parsec scale in the near- and mid-IR \citep[e.g.][]{Jaf04,Tri07,Bec08,Kis09a,Rab09,Tri09,Bur09}. The two prototypical type 2 AGN in NGC~1068 and the Circinus galaxy have been studied in detail using the mid-IR beam combiner MIDI at the Very Large Telescope Interferometer (VLTI) and pairwise combinations of the four 8\,m-telescopes. On these rather long baselines both objects were highly resolved with rather low visibilities that have been interpreted as the reemission signatures of the circumnuclear obscuring region, commonly dubbed ``dust torus''. This geometrically-thick torus is an essential part of the unification scheme of AGN and explains the difference of type 1 and type 2 AGN by angle-dependent obscuration. Recently, IR interferometry has shown that the spectral and spatial characteristics of the mid-IR emission sources (spectral slope and apparent size) in type 1 AGN are tightly correlated with the brightness profile and radial distribution of the dust \citep{Kis11b}. One problem with these previous observations is that they either miss most of the single-aperture flux, which has been resolved out (NGC~1068 and Circinus), or that they do not have the position-angle coverage necessary to constrain the shape/orientation of the nuclear emission source. This structural information is an important input for torus models that simulate IR images and predict certain position-angle dependent spectral and spatial features \citep{Sch08,Hon10b}. In this paper, we present mid-IR interferometry observations of the Seyfert 2 galaxy NGC~424. These observations cover almost the entire $uv$-plane accessible with the VLTI UTs. NGC~424 is a SB0/a galaxy at a distance of 45.7\,Mpc. The galactic disk is inclined at about 70$^\circ$ based on an axis ratio of 0.36 \citep{Kir90}. It hosts an AGN that is classified as an optical type 2 AGN with polarized broad lines \citep[i.e., a hidden type 1 AGN;][]{Ver10} based on spectro-polarimetric observations \citep{Mor00}. The gas column towards the nucleus is quite large: \citet{Col00} report that the Hydrogen column density $N_H$ is probably in the Compton-thick regime ($\ga2\times10^{24}\,\mathrm{cm}^{-2}$), while \citet{Lam11} find less, but still significant, X-ray obscuration ($N_H \sim 1.7\times 10^{23}\,\mathrm{cm}^{-2}$). Both values are consistent with a heavily obscured AGN, and most of the obscuring column is probably intrinsic to the AGN, in spite of the considerable inclination of the host galaxy. Based on the velocity dispersion of the bulge, \citet{Bia07} estimate the mass of the central black hole as $\log M_\mathrm{BH} = 7.78$, leading to an Eddington ratio of $l_\mathrm{Edd} = 0.13$, which is typical for a Seyfert galaxy. The nucleus is radio-quiet and unresolved in radio \citep{Mun00}. In Sect.~\ref{sec:obs}, we present the interferometric data and describe our data reduction strategy. For interpretation we also compiled a high spatial resolution IR SED of single-telescope data to complement the interferometry. In Sect.~\ref{sec:res}, wavelength- and position-angle-dependent sizes are extracted from the interferometric observations and the IR SED is analyzed. Both interferometry and photometry data are simultaneously modeled in Sect.~\ref{sec:model}. The results are discussed in Sect.~\ref{sec:disc} and put into context of AGN unification. Finally, we summarize our findings in Sect.~\ref{sec:summary}.
|
\label{sec:summary} We present VLTI/MIDI mid-IR interferometric observations of the X-ray-obscured type 2 nucleus of NGC~424. The $uv$-coverage allowed us to constrain the position angle-dependent sizes of the mid-IR emission source in the $8-13\,\micron$ wavelength range. We used two different methods to overcome size biases by the inhomogeneous distribution of projected baseline lengths on the $uv$-plane (i.e. beam shape effects). As complementary information, a high spatial resolution IR SED has been compiled. In addition to simple geometric fits, we also modeled photometric and interferometric data simultaneously by a simple model that assumes dust emission in thermal equilibrium with the AGN. This led to the following main results: \begin{itemize} \item The mid-IR emission source in NGC~424 is resolved and elongated. Its size depends on position-angle, wavelength, and baseline. For a baseline length of 58\,m we find a size of $(1.7\pm0.2)\,\mathrm{pc} \times (1.2\pm0.2)\,\mathrm{pc}$ at 9.1\,$\micron$, and $(2.2\pm0.4)\,\mathrm{pc} \times (1.9\pm0.4)\,\mathrm{pc}$ at 12.5\,$\micron$. As shown in Fig.~\ref{fig:res:wlsize}, the source size shows a systematic increase with wavelength. A more baseline-independent size can be extracted using the half-light radius $R_{1/2}$ as defined in \citet{Kis11b}. We obtain $R_{1/2} = (2.0\pm0.2)\,\mathrm{pc} \times (1.5\pm0.3)\,\mathrm{pc}$ for the $9-13\,\micron$ wavelength-averaged visibilities of NGC~424. \item The major axis of the mid-IR emission in NGC~424 points toward PA $-27^\circ\pm2^\circ$ based on the best-fit model of the complete data set. This direction is consistent with the PA derived from simple geometric fits to individual wavelength data corrected for the $uv$-coverage bias (or beam shape). We also detect a flatter radial brightness distribution in the direction of the major axis than along the minor axis of the mid-IR emission. Interestingly, the position angle of the major axis is only 20$^\circ$ off the system axis as set by spectro-polarimetry. Therefore, we conclude that the mid-IR emission is extended in polar direction or symmetry axis of the torus. \item The polar-elongated emission is responsible for $\ga$60\% of the mid-IR total flux (i.e. single telescope flux) in NGC~424 on scales from about 1\,pc to $\sim$100\,pc. \end{itemize} These results can be put into the broader context of AGN unification and our current picture of the mid-IR emission from the dusty torus. We conclude: \begin{itemize} \item The elongation of the nuclear mid-IR source on parsec scales in polar direction seems inconsistent with typical state-of-the-art radiative transfer models of the torus, either smooth or clumpy. While the hydrodynamic simulations by \citet{Sch09} do show polar-elongated mid-IR emission, the model SEDs are incompatible with the observed IR SED of NGC~424. We conclude that it is very difficult to simultaneously explain both the shape and SED of NGC~424 within the framework of current torus models. \item Based on the spectral (SED) and spatial (interferometry) information, the source of the emission is most likely optically thin dust in the polar region of the AGN. While the total dust column and surface brightness are, therefore, low, the emission covers a relatively large area leading to the high contribution of flux to the overall mid-IR emission. We note that similar mid-IR extensions in polar region have been observed on scales of several 10s of parsecs in a few other sources, including NGC~1068. \item The characteristics of the mid-IR emission in NGC~424 show similarities to the nucleus of the Circinus galaxy (and probably NGC~1068 as well) where about $\sim$85\% of the mid-IR flux originates in the polar region and coincides with the western edge of the outflow cone., i.e. several degrees off from the nominal system axis. In NGC~424, the mid-IR major axis is off about 20$^\circ$ from the polar axis. This could be interpreted as an edge brightening effect in a cone filled with optically thin dust. \item Our data suggests that the origin of the dust in the polar region may be a radiatively-driven wind from the inner part of the torus. The dust in the inner torus region is expected to be deficient of silicate grains, which is well matched by the mid-IR spectral features of the polar dust. Based on recent near- and mid-IR interferometry results of a sample of type 1 AGN, we propose that the torus has a small scale height at large distances from the AGN and is puffed-up by radiation pressure in the inner region, leading to the required covering factors in the unification scheme. In this picture, the near-IR emission (up to about $5\,\micron$ would be dominated by the torus, while the dusty outflow would be the main contributor to the mid-IR. \end{itemize} Our results for NGC~424 have significant implications for models of the IR emission of AGN. Owing to a lack of suitable data, we do not know if the mid-IR emission sources in NGC~424 and a few other nearby Seyfert galaxies are special cases or more generic to AGN, but it seems as more cases of polar-extended mid-IR emission are revealed when objects are studied in detail by high-angular resolution techniques. Moreover, the results of \citet{Hon10a} and \citet{Kis11b} suggest that the structure of the dusty environment depends on luminosity. It is, therefore, important to follow-up our results with interferometry from near- to mid-IR wavelengths and a good position-angle and baseline coverage of more sources. The upcoming VLTI/MATISSE instrument, operating from 3.5 to 13\,$\micron$, should be able to test if the near-IR emission is, indeed, dominated by the torus while the mid-IR flux originates in the outflow.
| 12
| 6
|
1206.4307
|
1206
|
1206.6905_arXiv.txt
|
Since the middle of the 1940's scientists have used Monte Carlo (MC) simulations to obtain information about physical processes. This has proved a accurate and and reliable method to obtain this information. Through out resent years researchers has begone to use the slightly newer Markov Chain Monte Carlo (MCMC) simulation. This differs from the ordinary MC by using the Markov Chain. MCMC originates from Bayesian statistics. This method has given researchers a completely new tool to learn something about physical systems. One of the fields where MCMC is a good new tool, is astrophysics. Today MCMC is widely used in simulating power spectra for asteroseismic data. Hereby providing the scientists with important new information of stellar interiors. From our results we see that MCMC delivers a robust and reliable result with good error estimation. We also learn that MCMC is a power full tool which can be applied to a large verity of problems.
| 12
| 6
|
1206.6905
|
||
1206
|
1206.6134_arXiv.txt
|
We estimate here a flux-transport dynamo model's response time to changes in meridional flow speed. Time-variation in meridional flow primarily determines the shape of a cycle in this class of dynamo models. In order to simultaneously predict the shape, amplitude and timing of a solar cycle by implementing an Ensemble Kalman Filter in the framework of Data Assimilation Research Testbed (DART), it is important to know the model's sensitivity to flow variation. Guided by observations we consider a smooth increase or decrease in meridional flow speed for a specified time (a few months to a few years), after which the flow speed comes back to the steady speed, and implement that time-varying meridional flow at different phases of solar cycle. We find that the model's response time to change in flow speed peaks at four to six months if the flow change lasts for one year. The longer the changed flow lasts, the longer the model takes to respond. Magnetic diffusivity has no influence in model's response to flow variation as long as the dynamo operates in the advection-dominated regime. Experiments with more complex flow variations indicate that the shape and amplitude of flow-perturbation have no influence in the estimate of model's response time.
|
There has been substantial interest in predicting future solar cycles for the past forty years \citep{ohl66,ohl79}. In the current era of extensive use of the high atmosphere and neighboring interplanetary medium by man, such predictions have considerable practical value. For cycles 22 and 23, prediction methods were primarily statistical rather than dynamical. That is, no physical laws were integrated forward in time, as is done for meteorological and climate predictions. But for solar cycle 24 the first such cycle prediction, which involves integrating forward in time a form of Faraday's law of electromagnetic theory, has been made \citep{ddg06,ccj07}. Given the step-by-step successes of different kinematic dynamo models, starting from convection zone dynamos \citep{stix76}, interface dynamos \citep{parker93}, up to flux-transport dynamos \citep{ws91,csd95,durney95, dc99,krs01,jetal08,gdd09} applied to the Sun, a set of kinematic flux-transport dynamo equations was chosen to numerically integrate forward in time \citep{dg06}. This was analogous to what was done in the 1950s with the earliest weather forecast models, which were 2D latitude-longitude models. Kinematic flux-transport dynamo models are also 2D, but in latitude and radius. The kinematic dynamo equations were calibrated with solar observations, and driven by input of observed solar magnetic data. The data input was continuous in time but quite simple -- a form of 'data-nudging', previously used in early weather and climate forecast models. These calculations simulated the relative peaks of the past cycles \citep{ddg06,ccj07} and showed skill even when North and South hemispheres were simulated separately \citep{dgdg07}. However, the predictive skill has been limited to hindcasting the peak-amplitude of a cycle in those calculations; for example, see Figure 1. Here a simulated sequence of cycles derived from a tachocline toroidal flux integral has been superimposed on observed cycles of monthly smoothed sunspot number (taken from www.sidc.be). The details within a cycle, such as its shape and its rise and fall patterns, have not been reproduced. In order to be able to predict the amplitude, timing and shape of a cycle simultaneously, we need to go beyond simple data-nudging for the entire span of integration. Updating the unknown time variations in the dynamo ingredients in a finite interval within a cycle, say every six to twelve months, will be required. Therefore, we need to implement a sophisticated data-assimilation scheme. \begin{figure}[hbt] \epsscale{1.0} \plotone{f1.eps} \caption{ Gray-filled curve shows observed cycles derived from monthly smoothed sunspot number data from the Royal Observatory of Belgium (www.sidc.be); superimposed on that is a simulated sequence of cycles (black curve) derived from total tachocline toroidal flux integral. Simple data-nudging for 12 successive cycles can hindcast the cycle-peaks with high skill, but not the details within a cycle. } \label{cycle-flux} \end{figure} Modern Earth system prediction models use sophisticated data assimilation methods to capture all the usable observational information about the system. These methods are highly developed for atmospheric and oceanic predictions \citep{kalnay03}, but have only recently begun to be used for the Sun \citep{brun07, kk09, jbt11}. With our present motivation of simulating cycle-shape and its rise and fall patterns, a suitable method is the Ensemble Kalman Filter (EnKF) technique in the framework of Data Assimilation Research Testbed (DART), which has been widely developed at the National Center for Atmospheric Research \citep{a01, awzh05, ac07, a09}. Identifying the parameters which govern the spatio-temporal pattern of meridional flow as the so-called state vectors, we can apply an Ensemble Kalman Filter to these state vectors and can create an ensemble of time-varying magnetic fields by advancing our dynamo model. A built-in Monte Carlo step within DART selects the simulation closest to observation after each specified time advancement of the model within a solar cycle. If we advance the model sequentially over an entire solar cycle and determine the rise and fall patterns of that cycle that match best with observations, then we can construct the spatio-temporal pattern of meridional flow for an entire cycle. Similarly, if we know the spatio-temporal variations of meridional flow in a cycle, we can predict the shape, rise and fall patterns of that cycle. \clearpage \begin{figure}[hbt] \epsscale{1.0} \plotone{f2.eps} \caption{ Thick gray curve represents an observed cycle based on monthly smoothed sunspot number; superimposed on it is a corresponding simulated cycle measured by its tachocline flux integral. Successive dashed, dotted and dash-triple-dotted curves describe a sequential data-assimilation method with three successive flow variations during the forward time-advancement of the dynamo model. Meridional flow-speed is adjusted in the model after each specified time-interval during a simulation-run. Combining the curves -- dashed, dotted and dash-triple-dotted curves and those beyond -- one member of the Ensemble can be constructed. } \label{schem-diagram} \end{figure} \clearpage Figure 2 describes how the cycle shape can change when the meridional flow speed changes after a specified time. From Figure 1, the observed cycle 22, derived from monthly smoothed sunspot number data, has been plotted in Figure 2 as a thick gray curve, and the simulated cycle (black curve) has been superimposed on it. As an example, Figure 2 clearly reveals that the simulated cycle's rise and fall patterns do not match with that of the observed one, because the simulation was performed with a steady meridional flow (see \citet{dg06} for details). Spatio-temporal variations in meridional flow were not known before the 1980's. Very recently the surface flow-patterns as functions of latitude and time are being detected \citep{ulrich10} (see also \citet{grbd11}). So, running the dynamo simulation, including updating the meridional flow speed after a specified time, and checking how the model-output compares with the observation, we can construct members of the Ensemble for a desired solar cycle (e.g. cycle 22, the example in Figure 2) and also the spatio-temporal variations of the meridional flow for that cycle. From our prior knowledge of the properties of flux-transport dynamos, we can make certain guesses about how the simulated cycle's phase will progress as the meridional flow varies. For example, near the beginning of the cycle (see Figure 2), we see that the simulated cycle-phase has progressed much faster than the observed one. Among various possibilities, a decrease in the flow-speed would make the cycle-phase progress more slowly, as shown by the dashed line. After a specified time, by comparing model-output with observations, the flow-speed can be updated again, with the aim of securing the closest possible match of the model cycle phase with the observed phase. In this regard, it is first necessary to know which ingredients in the model determine the shape of a cycle, how those ingredients vary with time and how the model responds to their time-variation. Previous flux-transport dynamo studies \citep{dc99} indicate that the meridional circulation is the key ingredient that determines the timing and shape of a cycle. We know from observations that the meridional circulations vary substantially with time for both the Earth and the Sun. Perturbations in meridional flow are likely to be particularly important, since from considerations of mass conservation these perturbations are likely to be felt rather quickly throughout the model domain. However, to make effective use of all the observational data available to predict future behavior of the Sun or Earth system, it is important to determine the 'response time' of these systems to perturbations of various types and durations occurring at various places within the system. This has been studied extensively for ocean circulation and climate systems \citep{neelin11, aa08, s04, srs06, sm99, hch80, bh89}, but not yet for the solar case. There does not appear to be a single 'response time', but rather a continuous range of times, beginning soon after the time the meridional circulation starts changing, continuing through a time of 'peak' response, followed thereafter by a time of declining further response. The time of peak response itself varies considerably, depending on the duration and location of the perturbation introduced, as well as the range of timescales inherent in the physics of the model. The model may respond quite differently to a perturbation of long duration compared to one of very short duration (for a discussion of this point in the case of climate systems, see \citet{neelin11} section 6.8). Comparison of the solar cycle prediction problem to that for the prediction of the Earth's climate system is particularly instructive. It is well known \citep{neelin11} that Earth's climate and its variations in time are determined by interactions among the various major components of the Earth 'system' (atmosphere, oceans, land surface, polar ice caps) as well as by external solar forcing. These interactions and forcings occur on a wide range of timescales, from days to centuries and millennia \citep{jbl00,dmc07,wdc00,ffsw06,jil11,cm05}. This is analogous to, but even more complex than, the Sun's 'climate' system, or the solar dynamo, in which timescales of days (for emergence of new magnetic flux in active regions) to years (for variations in the Sun's 'conveyor belt' or meridional circulation) to a decade or two (for transport of magnetic flux to the bottom of the convection zone) and beyond (for the envelope of the solar cycle and the occurrence of 'Maunder minima") are prominent. In the Earth system, with the exception of the surface mixing layer, the timescales in the ocean are much longer than those in the atmosphere. They range from months to years for the layer above the thermocline, to decades to millennia for the deeper ocean. It follows that to predict changes in climate on timescales longer than a few weeks, the dynamics and thermodynamics of the ocean must be included in the models that are used \citep{neelin11,czd86}. Analogously, to simulate solar cycle properties requires models that capture the MHD of the whole solar convection zone, for which the timescales are much longer than for emergence of new magnetic flux at the photosphere. Mean field flux-transport dynamo models are among the simplest that do this. In both the Earth and Sun meridional circulation plays a critical role in determining behavior of the respective systems on longer timescales \citep{w02,getal05,dc99,ddg06,dg06}. The closely related 'signal storage' capacity or memory in both systems is particularly important. In the ocean the memory of past temperature anomalies at the ocean-atmosphere interface is \citep{czd86} retained in the deeper ocean, brought there, and later back to the surface of the ocean, by meridional flow. Similarly, the memory of past photospheric magnetic flux patterns is retained deep in the convection zone, brought there by the predominantly inward meridional flow at high latitudes. This memory provides the basis for prediction of changes in the climate of the Earth and Sun on timescales of years to a decade or two. For the Earth, these predictions are focused on El Nino and La Nina events, as well as associated extratropical changes \citep{czd86,metal09}. For the Sun, the focus is on predicting how certain global properties of solar cycles, such as peak amplitude, duration and shape differ from one cycle to the next. This paper begins the process of developing a more sophisticated data assimilation scheme by estimating a flux-transport dynamo model's response time to variations in one of its crucial ingredients, the meridional flow. The results will be important for achieving skill in predicting details within a cycle, the cycle-shape, its rise and fall patterns.
|
Solar cycle prediction research has progressed significantly during recent years so that physics-based models are being integrated forward in time, in addition to using empirical relations. But so far those efforts have been limited to predictions of peak-amplitude of a cycle and its average duration. Attempts to predict simultaneously the shape, timing and amplitude of a cycle have not been made yet. With knowledge gained from advances in predictions employing oceanic and atmospheric climate models, it has proven necessary to move beyond simple data-nudging to more sophisticated data-assimilation techniques. Efforts using an Ensemble-Kalman Filter (EnKF) approach \citep{kk09} as well as variational approach \citep{jbt11} have been made using one-dimensional dynamo models. Prior results from flux-transport dynamo simulations reveal that the meridional circulation primarily governs the progress of a cycle's phase in this particular class of dynamo models. The time variations in meridional circulation speed and profile can be used in a sequential data-assimilation approach that involves an EnKF method in the framework of the Data Assimilation Research Testbed (DART) \citep{a09} in order to predict the details within a cycle, namely the rise and fall patterns, the onset, peak and end timings and the peak-amplitude. A sequential data assimilation approach can be most efficiently used if we know the model's response time to a change in its ingredients. Then we will know how to update the input data of a specific ingredient into the model. With this motivation we studied a flux-transport dynamo model's response time to meridional flow-speed variations, and found that the model's time of peak response to a change in flow-speed that lasts typically from one-half to one and a half years, is about six months on average. This response time is independent of proxies used to measure a theoretical solar cycle, such as tachocline toroidal field at $15^{\circ}$ latitude, at $60^{\circ}$ latitude or the integrated total toroidal flux in the tachocline. All response times are much shorter than the 'circulation time' or the time it takes for a fluid element to make a complete circuit on a closed streamline that reaches both low and high latitudes as well as passing close to the inner and outer radial boundaries of the dynamo domain. Incorporating changes in flow-speed lasting for different time-spans, such as 3, 6, 9, 12 and 18 months, we found that the time of peak response of the model increases with the duration of flow variation; it is approximately the length of full width at half maximum of the flow-perturbation profile in time. Response of the model is found to be always slightly faster when the flow change is positive with respect to the mean, steady flow, primarily because the rate of progress in a cycle's phase is approximately proportional to the flow-speed in this class of dynamo models. We also found that the time of peak response is independent of magnetic diffusivity so long as the dynamo operates in the advection-dominated regime. Although in most of our numerical experiments we chose a smooth sinusoidal type variation in flow-speed, we have also performed numerical experiments incorporating different shape and amplitudes of flow perturbations. We found that the model's time of peak response to change in flow is roughly independent of the shape and amplitude of the flow perturbations. There exists observations of systematic decrease of flow speed during the entire rising phase of the cycle 23 \citep{ba2003}. Consideration of the flow speed change during such a long span of time for studying the response time of a dynamo models is beyond the scope of this paper. In a sequential assimilation scheme the unknown model ingredients require updating more frequently than on a solar cycle time scale in a dynamo model that is attempting to simulate shape of a solar cycle. So the experiments we have performed here have the perturbation lasting for no more than two years. However, it will be interesting to investigate whether a dynamo model would respond within the same cycle or in the next cycle to a systematic flow variation that occurs during an entire rising or declining phase. We have focused only on the speed variation and taken no variation in latitudinal or radial flow profile, because we do not have information from observations about the complete flow-profiles in the convection zone. We can implement the knowledge gained here about the model's response time to change in flow-speed to develop EnKF schemes for assimilating time-varying flow data sequentially and simulate a cycle's rise and fall patterns along with its amplitude and timing. Flux-transport dynamo models are particularly sensitive to a meridional flow changes; so its influence may be overestimated in this class of models. Data assimilation techniques are useful to better determine the relevant ingredients in the mean-field dynamo models and quantify their influence on observables like the magnetic cycle period.
| 12
| 6
|
1206.6134
|
1206
|
1206.5906_arXiv.txt
|
We present observational data for two main components (S255IR and S255N) of the S255 high mass star forming region in continuum and molecular lines obtained at 1.3 mm and 1.1 mm with the SMA, at 1.3 cm with the VLA and at 23 and 50 cm with the GMRT. The angular resolution was from $ \sim 2'' $ to $ \sim 5'' $ for all instruments. With the SMA we detected a total of about 50 spectral lines of 20 different molecules (including isotopologues). About half of the lines and half of the species (in particular N$_2$H$^+$, SiO, C$^{34}$S, DCN, DNC, DCO$^+$, HC$_3$N, H$_2$CO, H$_2$CS, SO$_2$) have not been previously reported in S255IR and partly in S255N at high angular resolution. Our data reveal several new clumps in the S255IR and S255N areas by their millimeter wave continuum emission. Masses of these clumps are estimated at a few solar masses. The line widths greatly exceed expected thermal widths. These clumps have practically no association with NIR or radio continuum sources, implying a very early stage of evolution. At the same time, our SiO data indicate the presence of high-velocity outflows related to some of these clumps. In some cases, strong molecular emission at velocities of the quiescent gas has no detectable counterpart in the continuum. We discuss the main features of the distribution of NH$_3$, N$_2$H$^+$, and deuterated molecules. We estimate properties of decimeter wave radio continuum sources and their relationship with the molecular material.
|
Despite their great importance in almost all areas of astronomy, the formation of stars larger than 8--10 M$_\odot$ is still poorly understood. In part this is because high mass star formation (HMSF) regions are more distant, more active, and shorter lived than their low-mass counterparts. Nevertheless, an enhanced attention has been paid recently to the earliest phases of massive star formation. So far, few young HMSF cores have been studied in detail and it is important to extend the list of such objects, especially considering the large variety of processes that probably lead to the formation of a massive star. S255 is an \htwo\ region associated with a dense core at a commonly accepted distance of 2.5~kpc \citep{Russeil07,Ojha11} which we adopt here too. We note, however, that \citet{Rygl10} report a distance of 1.6~kpc based on trigonometric parallax measurements of methanol masers. The core consists of two main components (S255IR and S255N) separated by slightly over $1'$. Our team has previously acquired molecular line observations using single-dish instruments (OSO-20m, IRAM-30m, NRAO-12m). With angular resolution from about $1'$ to $10^{\prime\prime}$ \citep{Zin09}, we obtained $M \sim 300$~M$_{\odot}$, $n \sim 2 \cdot 10^5$~cm$^{-3}$, $T_{\mathrm{kin}} \sim 40$~K and $\Delta V \sim 2$~km\,s$^{-1}$ for both components. While both components show evidence for cool, massive clumps, their evolutionary states appear to be quite different. S255IR is bright ($> 70$ Jy) in all the Midcourse Space Experiment (MSX) bands and contains a near-IR cluster of early-B-type stars \citep{Howard97,Itoh01}, a cluster of compact \htwo\ regions \citep{Snell86}, and a wealth of complex H$_2$ emission features \citep{Miralles97}. In contrast, S255N (also called Sh2-255 FIR1 and G192.60-MM1) contains a single cometary UC \htwo\ region \citep[e.g.,][]{Kurtz94}, and is undetected by MSX at wavelengths shorter than 21~$\mu$m \citep{Crowther03}. VLA and SMA observations in the continuum, several molecular lines and water maser emission \citep{Cyganowski07} indicate the presence of a massive protocluster in this region. The chemical composition of S255IR and S255N also appear significantly different \citep{Lintott05,Zin09}. While the CS and HCN abundances are very similar, the abundances of NH$_3$, N$_2$H$^+$, HCO$^+$ and some other molecules in these components differ significantly. Recently, both components were studied with the SMA at 1.3 mm in CO, $^{13}$CO, C$^{18}$O, CH$_3$OH, CH$_3$CN and some other lines \citep{Wang11}. These observations revealed 3 continuum clumps in the S255IR area and high-velocity collimated outflows in both regions. The star population in this complex was studied most recently by \citet{Ojha11} on the basis of optical and NIR observations. They found a number of new YSO candidates with a large spread in ages, indicating a scenario of induced star formation. In general, this star forming complex represents an excellent laboratory for studies of different stages of massive star formation. Our goal is to investigate further the structure, physical properties and chemistry of this area on small scales by observations of important molecular tracers like NH$_3$ (which is a convenient ``thermometer" for dense molecular cores and provides information on their density and kinematics too --- e.g. \citealt{Ho83,Walmsley83}), N$_2$H$^+$, SiO, deuterated molecules, etc. In addition we want to investigate in more detail the distribution and properties of ionized gas in this region in order to trace the interaction of massive YSOs and outflows with the surrounding medium. In this paper we present observational data obtained at 1.3 mm and 1.1 mm with the SMA in the compact configuration (both in various lines and in continuum), with the VLA at 1.3 cm (in ammonia lines) and with the GMRT at 23 and 50 cm in continuum. We also discuss the general structure and kinematics of the complex and derive basic physical parameters of the observed features.
|
1. We detected several new clumps in the S255IR and S255N areas by their millimeter wave continuum emission. They are also detected in several molecular lines. Temperatures of these clumps found from ammonia observations are $ \sim 20 $~K and their masses are estimated at a few solar masses. These clumps have almost no association with NIR or radio continuum sources, which implies a very early stage of evolution. At the same time, our SiO and other molecular data indicate the presence of high-velocity outflows related to some of these clumps. The outflows associated with S255N-SMA3 and S255N-SMA5 are apparently very young --- a few hundred years only. The line widths of the clumps greatly exceed the expected thermal widths, suggesting significant turbulence in these objects. 2. In some cases there is strong molecular emission at the velocity of the quiescent gas, yet with no detected continuum counterpart. The nature of these sources is not clear. Some of them (e.g., the ammonia source in the northern part of S255N) may represent very cold gas not detectable in the dust continuum. Emission from CO and other molecules would be absent in this case, because they would be frozen onto dust grains. This assumption is consistent with the low kinetic temperature derived for this area from the ammonia observations. 3. N$_2$H$^+$ and DCO$^+$ apparently avoid ionized regions. There is no detectable N$_2$H$^+$ emission associated with the SMA1 clump in S255IR (which contains a hot core and also ionized gas) implying a significant drop in the N$_2$H$^+$ abundance. There is no sign that this is caused by CO evaporation. 4. We detected rather strong emission in lines of deuterated species --- DCN and DCO$^+$ --- in both S255IR and S255N (in S255N the DCN $ J=3-2 $ line was seen earlier by \citealt{Cyganowski07}). In S255N a DNC emission line is also detected. In S255IR the DCN emission is also quite strong in the vicinity of the hot core. This implies either the presence of a significant amount of cold gas or a very young age for this hot core, insufficient to have changed the isotopic ratio. The DCO$^+$ distribution is significantly different from that of DCN, and qualitatively resembles that of N$_2$H$^+$. In particular, both species avoid ionized regions, and both are observed towards the S255N SMA6 clump (in contrast to most other molecules). 5. There is no molecular material associated with the S255-2a and S255-2b radio continuum sources, which are compact \htwo\ regions excited by B type stars. Possibly the parental material of these stars has already been dispersed, indicating a relatively large age for these objects.
| 12
| 6
|
1206.5906
|
1206
|
1206.1862_arXiv.txt
|
Millisecond pulsars are mainly characterised by their spin periods, B-fields and masses -- quantities which are largely affected by previous interactions with a companion star in a binary system. In this paper, we investigate the formation mechanism of millisecond pulsars by considering the pulsar recycling process in both intermediate-mass {X}-ray binaries (IMXBs) and low-mass {X}-ray binaries (LMXBs). The IMXBs mainly lead to the formation of binary millisecond pulsars with a massive carbon-oxygen ({CO}) or an oxygen-neon-magnesium white dwarf ({ONeMg}~WD) companion, whereas the LMXBs form recycled pulsars with a helium white dwarf ({He}~WD) companion. We discuss the accretion physics leading to the spin-up line in the $P\dot{P}$--diagram and demonstrate that such a line cannot be uniquely defined. We derive a simple expression for the amount of accreted mass needed for any given pulsar to achieve its equilibrium spin and apply this to explain the observed differences of the spin distributions of recycled pulsars with different types of companions. From numerical calculations we present further evidence for significant loss of rotational energy in accreting {X}-ray millisecond pulsars in LMXBs during the Roche-lobe decoupling phase (Tauris~2012) and demonstrate that the same effect is negligible in IMXBs. We examine the recycling of pulsars with {CO}~WD companions via Case~BB Roche-lobe overflow (RLO) of naked helium stars in post common envelope binaries. We find that such pulsars typically accrete of the order $0.002-0.007\,M_{\odot}$ which is just about sufficient to explain their observed spin periods. We introduce isochrones of radio millisecond pulsars in the $P\dot{P}$--diagram to follow their spin evolution and discuss their true ages from comparison with observations. Finally, we apply our results of the spin-up process to complement our investigation of the massive pulsar \psr\ from Paper~I, confirming that this system formed via stable Case~A RLO in an IMXB and enabling us to put new constraints on the birth masses of a number of recycled pulsars.
|
\label{sec:Intro} Binary millisecond pulsars (BMSPs) represent the advanced phase of stellar evolution in close, interacting binaries. Their observed orbital and stellar properties are fossil records of their evolutionary history. Thus one can use binary pulsar systems as key probes of stellar astrophysics. It is well established that the neutron star component in binary millisecond pulsar systems forms first, descending from the initially more massive of the two binary zero-age main sequence (ZAMS) stars. The neutron star is subsequently spun up to a high spin frequency via accretion of mass and angular momentum once the secondary star evolves \citep{acrs82,rs82,bv91}. In this recycling phase the system is observable as a low-mass {X}-ray binary \citep[e.g.][]{hay85,nag89,bcc+97} and towards the end of this phase as an {X}-ray millisecond pulsar \citep{wv98,asr+09,pw12}. Although this standard formation scenario is commonly accepted many aspects of the mass-transfer process and the accretion physics are still not understood in detail \citep{lv06}. Examples of such ambiguities include the accretion disk structure, the disk-magnetosphere transition zone, the accretion efficiency, the decay of the surface B-field of the neutron star and the outcome of common envelope evolution. The current literature on the recycling process of pulsars is based on a somewhat simplified treatment of both the accretion process and the so-called equilibrium spin theory, as well as the application of a pure vacuum magnetic dipole model for estimating the surface B-field strength of a radio pulsar. These simplifications become a problem when trying to probe the formation and the evolution of observed recycled radio pulsars located near the classical spin-up line for Eddington accretion in the $P\dot{P}$--diagram \citep[e.g.][]{faa+11}. In this paper we discuss the concept and the location of the spin-up line and investigate to which extent it depends on the assumed details of the accretion disk/magnetosphere interactions and the magnetic inclination angle of the pulsar. Furthermore, we include the plasma filled magnetosphere description by \citet{spi06} for determining the surface B-field strengths of pulsars in application to spin-up theory and compare to the case of using the standard vacuum dipole model. A key quantity to understand the formation of any given recycled radio pulsar is its spin period which is strongly related to the amount of mass accreted. The amount of accumulated mass is again determined by the mass-transfer timescale of the progenitor {X}-ray binary (and partly by the mode of the mass transfer and mass loss from the system), and hence it depends crucially on the nature of the donor star. At present the scenario is qualitatively well understood where low-mass {X}-ray binaries (LMXBs) and intermediate-mass {X}-ray binaries (IMXBs) in general lead to the formation of pulsars with a helium white dwarf ({He}~WD) or a carbon-oxygen white dwarf ({CO}~WD) companion, respectively \citep[e.g.][]{tvs00,prp02,del08,tau11}. However, here we aim to quantify this picture in more detail by re-analysing the spin-up process. As an example, we investigate in this paper the possibilities of spinning up a neutron star from Case~BB RLO leading to a mildly recycled pulsar with a {CO}~WD companion. We also discuss the increasing number of systems which apparently do not fit the characteristics of these two main populations of binary pulsars. We follow the pulsar spin evolution during the final phase of the mass transfer by including calculations of the torque acting on an accreting pulsar during the Roche-lobe decoupling phase \citep[RLDP;][]{tau12} for both an LMXB and an IMXB and compare the results. To complete the full history of pulsar spin evolution we subsequently consider the true ages of recycled radio pulsars (which are intimately related to their spin evolution) by calculating isochrones and discussing the distribution of recycled pulsars in the $P\dot{P}$--diagram. Although accreting {X}-ray millisecond pulsars (AXMSPs) are believed to be progenitors of BMSPs, all of the 14 observed AXMSPs have orbital periods less than one day whereas fully recycled radio BMSPs are observed with orbital periods all the way up to a few hundred days \citep[see also][]{tau12}. This puzzle will be investigated in a future paper. For the evolution of ultra-compact {X}-ray binaries (UCXBs) leading to AXMSPs, and later to the formation of tight-orbit recycled radio BMSPs with ultra-light ($<0.08\,M_{\odot}$) companions, we refer to \citet{prp02,db03,vvp05,vnv+12}. In this paper we focus on the formation and evolution of BMSPs with regular {He}~WDs and {CO}/{ONeMg}~WDs. The discovery of \psr\ \citep{hrr+05,dpr+10} plays an important role for understanding BMSP formation. This pulsar system is interesting since it has a unique combination of a short pulsar spin period of 3.2~ms and a massive {CO}~WD companion. A rapidly spinning pulsar is usually associated with a long phase of mass-transfer evolution in an LMXB, whereas a {CO}~WD companion in a relatively close orbit (8.7~days) is evidently the outcome of an IMXB evolution. A possibly solution to this paradox is that \psr\ evolved from Case~A RLO of an IMXB which results in both a relatively long-lived {X}-ray phase ($>10^7\,{\rm yr}$), needed to spin~up the pulsar effectively, and leaving behind a {CO}~WD. In this case \psr\ is the first BMSP known to have evolved via this path. Indeed, in \citet{tlk11}, hereafter Paper~I, we investigated the progenitor evolution of \psr\ with emphasis on the {X}-ray phase where the binary contained a neutron star and a donor star. We found two viable possibilities for the formation of the \psr\ system: either it contained a $2.2-2.6\,M_{\odot}$ asymptotic giant branch donor star and evolved through a common envelope and spiral-in phase initiated by Case~C RLO, or it descended from a close binary system with a $4.0-5.0\,M_{\odot}$ main sequence donor star via Case~A RLO as hinted above. The latter scenario was also found by \citet{lrp+11}. The fact that \psr\ was spun-up to (less than) 3.2~ms could indeed hint which one of the two formation scenarios is most likely. In order to test this idea and to further distinguish between Case~A and Case~C we turn our attention, here in Paper~II, to the spin dynamics of this pulsar in the recycling process. As discussed in Paper~I, the distribution of neutron star birth masses is an important probe of both stellar evolution, binary interactions and explosion physics. For a number of BMSPs we are now able to put constraints on the birth mass of the pulsar given the derived amount of mass needed to spin up the observed recycled pulsar. In particular, it is of interest to see if we can identify further pulsars showing evidence of being born massive ($\sim\!1.7\,M_{\odot}$) like \psr. In order to understand the many different observational properties of BMSPs we have combined a variety of subtopics here with the aim of presenting a full picture of the subject. Given the many facets included, our new findings and the resulting length of the manuscript, we have chosen throughout this paper to finalize individual subtopics with a direct comparison to observational data followed by discussions in view of our theoretical modelling. Our paper is structured in the follow way: We begin with a brief, updated review of the formation channels of BMSPs with {CO}~WDs (Section~\ref{sec:formation}) and present a summary of the latest observational data in Section~\ref{sec:observations}. In this section we also demonstrate an emerging unified picture of pulsar formation history which, however, is challenged by a number of interesting systems which share expected properties of both an LMXB and an IMXB origin. In Section~\ref{sec:spinup} we investigate the recycling process in general with a focus on the location of the spin-up line in the $P\dot{P}$--diagram and also relate the initial spin of a rejuvenated pulsar to the amount of mass accreted. The theoretical modelling is continued in Section~\ref{sec:RLDP} where we highlight the effects of the Roche-lobe decoupling phase on the spin evolution of recycled pulsars. In Section~\ref{sec:obs-spin} our results are compared to the observed spin period distributions and in Section~\ref{sec:CEspin} we investigate if BMSPs with {CO}~WD companions obtained their fast spin periods {\it after} a common envelope evolution. In Section~\ref{sec:trueages} we discuss our results in a broader context in relation to the spin evolution and the true ages of millisecond radio pulsars. In Section~\ref{sec:1614-2230} we continue our discussion from Paper~I on the formation and evolution of \psr\ and in Section~\ref{sec:NSmass} we return to the question of neutron star birth masses. Our conclusions are summarized in Section~\ref{sec:summary}. Finally, in the Appendix we present a new tool to identify the most likely nature of the companion star to any observed binary pulsar.
|
\label{sec:summary} We have investigated in detail the recycling process of pulsars with respect to accretion, spin~up and the Roche-lobe decoupling phase, and discussed the implications for their spin periods, masses and true ages. In particular, we have discussed the concept of a spin-up line in the $P\dot{P}$--diagram and emphasize that such a line cannot be uniquely defined. Besides from the poorly known disk-magnetosphere physics, which introduces large uncertainties, for each individual pulsar the equilibrium spin period, $P_{\rm eq}$ also depends on its magnetic inclination angle, $\alpha$ as well as its accretion history ($\dot{M}$) and its B-field strength. Furthermore, we have applied the \citet{spi06} spin-down torque on radio pulsars which significantly changes the location of the spin-up lines compared to using the vacuum magnetic dipole model, especially for small values of $\alpha$. We have derived a simple analytical expression (equation~\ref{eq:deltaMfinalfit}) to evaluate the amount of mass needed to be accreted to spin~up a pulsar to any given equilibrium spin period, $P_{\rm eq}$. Our result resembles that of \citet{acrs82} and approximately yields the same values (within a factor of two) as the expression derived by \citet{lp84}. Using our formula we find, for example, that BMSPs with {He}~WDs and $P_{\rm eq}\simeq2\,{\rm ms}$ must have accreted at least $0.10\,M_{\odot}$, whereas typical BSMPs with {CO}~WDs and $P_{\rm eq}\simeq20\,{\rm ms}$ only needed to accrete $0.005\,M_{\odot}$. Applying equation~(\ref{eq:deltaMfinalfit}) enables us to explain the difference in spin distributions between BMSPs with {He} and {CO}~WDs, respectively. The BMSPs with {He}~WDs often evolved via an {X}-ray phase with stable RLO on a long timescale, allowing sufficient material to be accreted by the neutron star to spin it up efficiently to a short period, whereas the BMSPs with {CO}~WDs most often have slow spins and evolved from IMXBs which had a short phase of mass transfer via early Case~B RLO or Case~C RLO -- the latter leading to a CE-evolution often followed by Case~BB RLO. The only exception known so far is \psr\ since this system produced a {CO}~WD orbiting a fully recycled MSP and thus it must have evolved via Case~A RLO of an IMXB. It is not possible to recycle pulsars to become MSPs via wind accretion from $1.1-2.2\,M_{\odot}$ post-CE helium stars. However, we have demonstrated that Case~BB RLO from such helium stars can spin~up pulsars to at least $\sim\!11~{\rm ms}$. Further studies of these systems are needed. There is an increasing number of recycled pulsars with WD companions which seem to fall outside the two main populations of BMSPs with He and CO~WDs, respectively. These peculiar systems possibly have He~WDs and always exhibit slow spin periods between 10 and 100~ms. We suggest that these systems with $P_{\rm orb}\ge 1\;{\rm day}$ may have formed via Case~A RLO of IMXBs. We plan further studies on these binaries. The Roche-lobe decoupling phase (RLDP), at the terminal stage of the mass transfer \citep{tau12}, has been analysed and we have shown that while the RLDP~effect is important in LMXBs -- leading to significant loss of rotational energy of the recycled pulsars as well as characteristic ages at birth which may exceed the age of the Universe -- it is not significant in IMXB systems where the duration of the RLDP is short. In order to track the evolution of pulsars in the $P\dot{P}$--diagram we have introduced two types of true age isochrones -- one which matches well with the banana shape of the observed distribution of known MSPs. We encourage further statistical population studies to better understand the formation and evolution of radio MSPs in the $P\dot{P}$--diagram \citep[see e.g.][]{kt10}. The discrepancy between true ages and characteristic spin-down ages of recycled pulsars has been discussed and we confirm that the latter values are completely untrustworthy as true age indicators, leaving WD cooling ages as the only valid, although not accurate, measuring scale \citep{tau12}. In the combined study presented here and in Paper~I we have investigated the recycling of \psr\ by detailed modelling of the mass exchanging {X}-ray phase of the progenitor system. Given the rapid spin of \psr\ (3.15~ms) we argue that it is highly unlikely that it evolved through a CE, leaving Case~A RLO in an IMXB as the only viable formation channel. We confirm the conclusion from Paper~I that the neutron star in \psr\ was born massive ($1.70\pm0.15\,M_{\odot}$). We have demonstrated that \psr\ could have been spun-up at $\dot{M}=\dot{M}_{\rm Edd}$ and subsequently evolve to its current position in the $P\dot{P}$--diagram within 2~Gyr (the estimated cooling age of its white dwarf companion). Besides \psr, Vela~{X}-1 and possibly the Black-Widow pulsar, we have argued that also PSR~J0621+1002 could belong to the same class of neutron stars born massive ($\ge 1.7\,M_{\odot}$). The formation of such massive neutron stars in supernovae is in agreement with some supernova explosion models \citep[e.g.][]{zwh08,ujma12}.
| 12
| 6
|
1206.1862
|
1206
|
1206.6029_arXiv.txt
|
We argue that one of the basic assumptions of the Bondi accretion process, that the accreting object has zero pressure, might not hold in many galaxies because of the pressure exerted by stellar winds of star orbiting the central super massive black hole (SMBH). Hence, the Bondi accretion cannot be used in these cases, such as in the galaxy NGC~3115. The winds of these high-velocity stars are shocked to temperatures above the virial temperature of the galaxy, leading to the formation of a hot bubble of size $\sim 0.1-10 \pc$ near the center. This hot bubble can substantially reduce the mass accretion rate by the SMBH. If the density of the hot bubble is lower than that of the interstellar medium (ISM), a density-inversion layer is formed. {{{ As the gas loses energy by X-ray radiation, eventually more mass of the cooling shocked stellar winds will be accreted to the SMBH. This accretion will be of cold clumps. After a period of millions of years of low AGN activity, therefore, a stronger AGN activity will occur that will heat and expel gas, much as in cooling flow clusters. }}} Adding to other problems of the Bondi process, our results render the Bondi accretion irrelevant for AGN feedback in cooling flow in galaxies and small groups of galaxies and during galaxy formation.
|
\label{s-intro} It is widely accepted that feedback powered by active galactic nuclei (AGN) has a key role in galaxy formation and in cooling flows in galaxies and in clusters of galaxies. In galaxy formation AGN feedback heats and expels gas {{{ from the galaxy}}} (e.g., \citealt{Bower2008, Ostriker2010} and references therein), and by that can determine the correlation between the central super-massive black hole (SMBH) mass and some properties of the galaxy {{{ \citep{King2003,King2005,Soker2009b,Soker2011}}}}. In cooling flow clusters jets launched by the SMBH heat the gas and maintain a small, but non zero cooling flow (see review by \citealt{McNamara2007, McNamara2012, Fabian2012}); this is termed a moderate cooling flow. There is a dispute on how the accretion onto the SMBH occurs, in particular in cooling flows. One camp argues for accretion to be of hot gas via the Bondi accretion process (e.g., \citealp{Allen2006, Russell2010, Narayan2011}), while the other side argues that the accretion is of dense and cold clumps in what is termed the cold feedback mechanism {{{ \citep{Pizzolato2005,Pizzolato2010}}}}. The cold feedback mechanism has been strengthened recently by observations of cold gas and by more detailed studies \citep{Revaz2008,Pope2009,Wilman2009,Wilman2011,Nesvadba2011,Cavagnolo2011,Gaspari2012a,Gaspari2012b, McCourt2012,Sharma2012,Farage2012,Kashi2012}. The Bondi accretion process, on the other hand, suffers from two problems. The first problem is that in cooling-flow clusters the Bondi accretion rate is too low to account for the AGN power (e.g., \citealt{McNamara2011, Cavagnolo2011}). The second is that there is no time for the feedback to work \citep{Soker2009}. {{{ This is because the time for cooling gas at distances of $\ga {\rm few} \times \kpc$ in the Bondi accretion process to be accreted and power jets that heat back the ISM, is much longer than the cooling time of the gas. }}} This is already true for gas cooling at a moderate distance of $\sim 1 \kpc$ from the center. In other words the gas at large distances has no time to communicate with the SMBH before it cools. In this paper we point out yet another problematic point with the Bondi accretion process. In a recent paper, \citet{Wong2011} resolved the region within the Bondi accretion radius of the S0 galaxy NGC~3115. If the density and temperature profile is interpreted as resulting from a Bondi accretion flow onto the $M_{\rm BH}=2 \times 10^9 M_\odot$ central SMBH, the derived accretion rate is $\dot M_B=2.2 \times 10^{-2} M_\odot \yr^{-1}$. They note that for a radiation power of $0.1 \dot{M_B}\, c^2$, the expected accretion luminosity is six orders of magnitude above the observed upper limit. They attribute this to a process where most of the inflowing gas is blown away, {{{ or the gas is continuously circulating in convective eddies, or to that the region they resolve is not yet incorporated to the Bondi accretion flow. The idea of circulating eddies has some similarities to the density inversion layer behavior we discuss here. }}} {{{ In any case, some AGN activity does take place in NGC~3115 \citep{Wrobel2012}. \cite{Wrobel2012} detected a radio nucleus in NGC~3115 with a radio power of $L_{\rm radio}=3 \times 10^{35} \erg \s^{-1}$. This indicates the presence of a weak AGN, that might substantially reduce the accretion rate \citep{Wrobel2012}. As we discuss later, the feeding of the SMBH might be from the stellar winds rather than from the ISM. }}} {{{ Several other processes were considered to reduce the accretion rate by a SMBH much below the Bondi accretion rate. Such processes include magnetic field reconnection \citep{Igumenshchev2002}, angular momentum \citep{Proga2003a,Proga2003b}, magneto-thermal instabilities \citep{Sharma2008}, and instabilities due to self-gravitation of the infalling gas \citep{Levine2010}. Lack of spherical symmetry in realistic situations is an additional factor \citep{Debuhr2011}. Turbulent media can have higher than Bondi-Hoyle accretion rate, but due to vorticity, a lower accretion rate is also possible \citep{Krumholz2005,Krumholz2006}. \cite{Hobbs2012} claim that the Bondi-Hoyle solution is only relevant for hot virialized gas with no angular momentum and negligible radiative cooling. }}} We take a different view {{{ on the suppression of the Bondi accretion. We argue that in many galaxies for a fraction of the time the Bondi accretion flow might not be relevant because }}} one cannot assume a zero pressure at the center, either because of stellar winds or because of jets blown by the AGN.
|
\label{s-summary} We studied the pressure exerted by the winds of circum-SMBH high-velocity stars on the surrounding ISM. We found that in some cases this pressure is significant and can substantially suppress the inflow of the ISM relative to what a simple Bondi accretion would give. Our result can explain the finding of \citet{Wong2011} that the Bondi accretion rate calculated by them from the ISM density and temperature is six orders of magnitude above the observed upper limit on the accretion rate in the S0 galaxy NGC~3115. In section \ref{s-ngc3115} we quantitatively examined the situation in the galaxy NGC~3115. Shocked winds of circum-SMBH high-velocity stars form a bubble of hot gas whose pressure is significant, as evident from Fig. \ref{fig:Pr2}. {{{ The colliding winds heat up to very high temperatures, build significant pressure, and are not expected to be accreted by the SMBH even though they lose angular momentum. Cooler clumps that fall inward, from the ISM or from inhomogeneities within the hot bubble, will encounter the winds of fast-moving stars very close to the SMBH. This collision will heat such clumps, suppressing their accretion. Even if there is a small accretion rate, a very weak disc wind from the accretion disc might further lower the accretion rate. The study of the interaction of AGN winds with the gas near the SMBH is a subject of a future study using numerical simulations. }}} There are some uncertainties in the model, such as the exact behavior of the stellar mass loss, trajectories of stars around the SMBH, and the stochastic behavior of the post-shock stellar winds. Some of these will be studied in future numerical simulations. However, the result that the stellar winds cannot be ignored is robust. For some values of the parameters we found that a situation might arise where the hot bubble's density is lower than the ISM density. In this case, Rayleigh-Taylor (RT) instability takes place, and a density-inversion layer is formed (see schematic description in Fig. \ref{fig:fig1}). Although hot tenuous gas buoys outward and dense ISM gas moves inward, the density-inversion layer itself continues to exist. The ISM gas is heated near the center and accumulated {{{ into}}} the hot bubble. {{{ While the scenario suggested here may explain the low X-ray luminosity observed in the galaxy NGC~3115, its properties have not yet been observed or affirmed directly. The size of the hot bubble described is below the resolution limit of the observations and cannot yet be observed. Alternative explanations for a below-Bondi accretion rate are mentioned in section \ref{s-intro}. }}} {{{ We note that in our scenario there can be no steady state over a very long time of $ \sim 10^7-10^8 \yr$. Over this time scale radiative cooling becomes important and more of the cooling gas will be accreted by the SMBH. This will lead to stronger AGN activity that will heat and expel gas, hence reducing back the accretion rate and AGN power. In addition stellar formation must occur from time to time. Most likely, there are local star-burst episodes when the accretion rate is much higher than the Bondi accretion rate. The high accretion rate is probably driven by cold clumps (filaments, streams). Indeed, the stellar-wind pressure cannot prevent accretion of very dense clouds.}}} Our result is more general in showing that in many cases the Bondi accretion process does not work because one of its basic assumptions, that there is no central pressure, breaks down. This is one of several reasons why the Bondi accretion model may not apply in some cases (see section \ref{s-intro}). Finally, we note that our model may be relevant for active galaxies where the hot bubble might be formed by the AGN jets or winds. For typical values of AGN jets and winds the hot bubble density will be low, and a density-inversion layer will be formed. We expect this process to be of high significance in the process of AGN feedback acting in young galaxies. Barring Bondi-like accretion, dense and cold clumps in the ISM can still flow inward and feed the SMBH. Namely, AGN feedback mechanisms require the feeding to be by cold clumps, i.e., a cold feedback mechanism. {{{ We thank an anonymous referee for many detail and very helpful comments that substantially improved the manuscript. }}} This research was supported by the Asher Fund for Space Research and the E. and J. Bishop Research Fund at the Technion, and the Israel Science foundation.
| 12
| 6
|
1206.6029
|
1206
|
1206.4730_arXiv.txt
|
We present 230 GHz Submillimeter Array continuum and molecular line observations of the newly discovered FUor candidate HBC722. We report the detection of seven 1.3 mm continuum sources in the vicinity of HBC722, none of which correspond to HBC722 itself. We compile infrared and submillimeter continuum photometry of each source from previous studies and conclude that three are Class 0 embedded protostars, one is a Class I embedded protostar, one is a Class I/II transition object, and two are either starless cores or very young, very low luminosity protostars or first hydrostatic cores. We detect a northwest-southeast outflow, consistent with the previous detection of such an outflow in low-resolution, single-dish observations, and note that its axis may be precessing. We show that this outflow is centered on and driven by one of the nearby Class 0 sources rather than HBC722, and find no conclusive evidence that HBC722 itself is driving an outflow. The non-detection of HBC722 in the 1.3 mm continuum observations suggests an upper limit of 0.02 \msun\ for the mass of the circumstellar disk. This limit is consistent with typical T Tauri disks and with a disk that provides sufficient mass to power the burst.
|
FU Orionis objects (hereafter FUors) are a group of young, pre-main sequence stars observed to flare in brightness by $4-6$ magnitudes in the optical and remain bright for decades (Herbig 1977). They are named after the prototype FU Orionis, which flared by about 6 magnitudes in 1936 and has remained in an elevated state to the present day (Wachmann 1954; Herbig 1966). Only 10 confirmed FUors are known to exist from direct observations of flares, with about another 10 identified based on similar spectral characteristics to the confirmed FUors (see Reipurth \& Aspin 2010 for a recent review). The large amplitude flares are attributed to enhanced accretion from the surrounding circumstellar disk (Hartmann \& Kenyon 1985), with the accretion rate from the disk onto the star increasing to up to $\sim 10^{-4}$ \msun\ yr$^{-1}$ (Hartmann \& Kenyon 1996). Various triggering mechanisms for the accretion bursts have been proposed, including interactions with binary companions (Bonnell \& Bastien 1992), thermal instabilities (Hartmann \& Kenyon 1996), and gravitational and magnetorotational instabilities (Zhu et al.~2007, 2009a, 2009b; Vorobyov \& Basu 2010). FUors are especially interesting and relevant to general star formation studies because they may represent the late, optically visible end stages of episodic accretion bursts and luminosity flares through the duration of the embedded phase (e.g., Kenyon et al.~1990; Enoch et al.~2009; Evans et al.~2009; Vorobyov 2009; Dunham et al.~2010; Dunham \& Vorobyov 2012). With so few bona-fide FUors, detailed observations of each one are necessary in order to characterize their properties and understand their place in the general star formation process. In this paper, we present 230 GHz Submillimeter Array (SMA; Ho et al.~2004) observations of the newly discovered FUor candidate HBC722. As described in more detail in \S \ref{sec_hbc722} below, HBC722 is located within a small group of $\sim$ 10 young stars, greatly complicating the analysis of existing, low-resolution single-dish ground and space-based (sub)millimeter data. The SMA 230 GHz continuum and molecular line observations presented in this paper are motivated by a need to disentagle the millimeter emission from the various sources in the vicinity of HBC722 in order to better determine its evolutionary status and physical properties. The organization of this paper is as follows: a brief summary of HBC722 is given in \S \ref{sec_hbc722}, a description of the observations and data reduction is provided in \S \ref{sec_obs}, the basic results are presented in \S \ref{sec_results}, including the continuum data in \S \ref{sec_results_continuum} and the CO line data in \S \ref{sec_results_co}, a discussion of the detected continuum sources is presented in \S \ref{sec_discussion_sources}, a discussion of the evolutionary status of HBC722 is given in \ref{sec_discussion_hbc722}, and a summary of our results is presented in \S \ref{sec_summary}.
|
\label{sec_discussion} \subsection{Continuum Sources}\label{sec_discussion_sources} Figure \ref{fig_cont_overlay} displays infrared and submillimeter images of the HBC722 environment, including an UKIDSS (UKIRT [United Kingdom Infrared Telescope] Infrared Deep Sky Survey) $K_{\rm S}$ image, a \emph{Spitzer} 8 \um\ image from Guieu et al.~(2009), a \emph{Spitzer} 24 \um\ image from Rebull et al.~(2011), and a \emph{Herschel} 70 \um\ image, SHARC-II 350 \um\ image, and \emph{Herschel} 500 \um\ image from Green et al.~(2011). Overplotted in black are the SMA 1.3 mm continuum intensity contours and the primary beams of the SMA pointings. A version of this Figure without the SMA 1.3 mm continuum intensity contours was previously presented by Green et al.~(2011). Inspection of Figure \ref{fig_cont_overlay} shows that most of the seven detected SMA continuum sources are associated with sources at other wavelengths. A detailed discussion of each individual continuum source and its associations with sources at other wavelengths is given below. Table \ref{tab_photometry} lists, for each source, \emph{Spitzer} photometry at $3.6-24$ \um\ from Guieu et al.~(2009) and Rebull et al.~(2011), \emph{Herschel} 70 and 100 \um\ photometry from Green et al.~(2011), SHARC-II 350 \um\ photometry from Green et al.~(2011), and SMA 1.3 mm continuum photometry from this work. Figure \ref{fig_seds} plots SEDs of all seven continuum sources, including both detections and upper limits (see below for more details). \input tab3.tex \begin{figure*} \epsscale{1.0} \plotone{f7.eps} \caption{\label{fig_seds}Spectral Energy Distributions (SEDs) for each of the seven detected continuum sources, consisting of \emph{Spitzer} $3.6-24$ \um\ data, \emph{Herschel} 70 and 100 \um\ data, SHARC-II 350 \um\ data, and the SMA 1.3 mm continuum data (see text for details). Detections are plotted as filled circles with error bars and upper limits are plotted as open triangles. Each panel is labeled with the corresponding source.} \end{figure*} \input tab4.tex Table \ref{tab_evolindic} presents three evolutionary indicators for each source calculated from the SEDs tabulated in Table \ref{tab_photometry} and plotted in Figure \ref{fig_seds}: the infrared spectral index ($\alpha$), the bolometric temperature (\tbol), and the bolometric luminosity (\lbol). As first defined by Lada \& Wilking (1984) and Lada (1987), $\alpha$ is the infrared slope in log space of $\nu S_{\nu}$ vs.~$\nu$ and is used to classify sources into different evolutionary stages (see Evans et al.~[2009] for a recent review of classification via $\alpha$). In this study we calculate $\alpha$ with a linear least-squares fit to all available \emph{Spitzer} photometry between $3.6-24$ \um. \tbol\ is defined as the temperature of a blackbody with the same flux-weighted mean frequency as the source SED (Myers \& Ladd 1993) and provides an alternative classification method to $\alpha$ (Chen et al.~1995; Evans et al.~2009). We calculate both \tbol\ and \lbol\ by using the trapezoid rule to integrate over the finely sampled SEDs; a detailed description of the implementation of this method and resulting errors due to the finite sampling of the observed SEDs is given in Appendix B of Dunham et al.~(2008). An additional error is introduced by the fact that the SMA 1.3 mm continuum data resolve out some of the emission from the extended core and are thus lower limits to the true flux densities at this wavelength, artifically steepening the far-infrared and submillimeter slope of the SEDs. The magnitude of the error introduced depends on both the amount of emission recovered by the SMA and the spectral shape of each source; a very conservative estimate that only 1\% of the emission is recovered leads to underestimates in \lbol\ by less than a factor of 2 for all sources, and less than 50\% for all but MMS7. Thus, while we caution that the calculated \lbol\ may underestimate the true values, the magnitude of these underestimates are likely comparable to or less than the other errors discussed by Dunham et al.~(2008). Finally, we note that there are three other infrared sources identified as \emph{Spitzer}-detected Young Stellar Objects by Guieu et al.~(2009) and Rebull et al.~(2011) covered by our SMA observations, and several other infrared sources covered that are not identified as Young Stellar Objects and are thus likely background or foreground stars. As our focus is on providing a high spatial resolution view of the millimeter emission in the vicinity of HBC722 we do not discuss these sources further except to note that their non-detections in our SMA 1.3 mm continuum observations suggest similar upper limits to the masses of any circumstellar disks surrounding these objects as for HBC722 itself, which we discuss below in \S \ref{sec_discussion_hbc722}. \subsubsection{MMS1} MMS1 is associated with the \emph{Spitzer} infrared source SST J205816.56$+$435352.9 from Guieu et al.~(2009) and Rebull et al.~(2011). It is detected by \emph{Spitzer} only at 24 \um\ with the flux density listed in Table \ref{tab_photometry}. The \emph{Spitzer} $3.6-8$ \um\ upper limits listed are taken from Guieu et al.~(2009); they note that their 90\% completeness limits increase from about 0.3 mJy at 3.6 \um\ to 0.6 mJy at 8 \um\ but do not specifically list these limits for 4.5 and 5.8 \um, thus we conservatively take the upper limits to be 0.6 mJy in all four bands. MMS1 is also associated with a \emph{Herschel} source detected at 70 and 100 \um\ from Green et al.~(2009); all other \emph{Herschel} wavelengths are of too low spatial resolution ($18-35$\as\ at $160-500$ \um) to accurately separate the 7 continuum sources and are not considered in this study. The 70 and 100 \um\ photometry presented in Table \ref{tab_photometry} is calculated with 10\as\ diameter apertures chosen as the best compromise between including as much source flux as possible and excluding flux from other, nearby sources. No aperture or color corrections are applied since neither the true spatial profile of the emission nor the underlying spectral shape of any source is known. Finally, MMS1 is also associated with a SHARC-II 350 \um\ source. The photometry presented in Table \ref{tab_photometry} is calculated in a 20\as\ diameter aperture following the method described by Wu et al.~(2007). The SED of MMS1 resembles that of a deeply embedded protostar. We are unable to calculate $\alpha$ with only one \emph{Spitzer} detection, but the upper limits are consistent with a positive $\alpha$ indicative of a Class 0/I object. We calculate \tbol\ $=39$ K, placing MMS1 in Class 0, consistent with the above statements. We thus conclude that MMS1 is a deeply embedded Class 0 protostar. As described in \S \ref{sec_results_co} above, we do not detect clear signatures of an outflow driven by MMS1 but do note that there may be additional outflows in this region that are not fully separated spatially and kinematically by these data. \subsubsection{MMS2} MMS2 is associated with the \emph{Spitzer} infrared source SST J205816.8$+$435335.6 from Guieu et al.~(2009) and Rebull et al.~(2011) and is detected at all five \emph{Spitzer} wavelengths from $3.6-24$ \um. It is also detected at 70 and 100 \um\ with \emph{Herschel}. The photometry is again calculated in 10\as\ diameter apertures and is uncertain since the apertures do not fully capture the emission that extends to the northwest but also partially overlap with the emission from the brighter MMS3. More accurate photometry at these wavelengths will require higher-resolution observations. There is no clear SHARC-II 350 \um\ source, but the position of MMS2 overlaps with bright emission from MMS3 and MMS5. We calculate the upper limit as the flux in one beam at the position of MMS2 from the other, nearby sources. The SED of MMS2 resembles that of a Class I protostar more evolved and less deeply embedded than MMS1. The calculated values of $\alpha$ and \tbol\ (0.49 and 147 K, respectively) both classify MMS2 as Class I, confirming this statement. \subsubsection{MMS3} MMS3 is associated with the \emph{Spitzer} infrared source SST J205817.7$+$435331.1 from Guieu et al.~(2009) and Rebull et al.~(2011) and is detected at $3.6-24$ \um\ with \emph{Spitzer}. It is also detected at 70 and 100 \um\ with \emph{Herschel} and 350 \um\ with SHARC-II and is the brightest source in the region at these wavelengths. The photometry at 70, 100, and 350 \um\ is calculated with apertures and methods identical to MMS1 above. The true flux densities at these wavelengths are likely higher than the values listed in Table \ref{tab_photometry} since the apertures do not include all of the extended emission, but larger apertures are not feasible since they would overlap with other, nearby sources. As with MMS1, the SED of MMS3 resembles that of a deeply embedded protostar. The calculated values of both $\alpha$ (0.76; in the Class 0/I category) and \tbol\ (52 K; in the Class 0 category) are consistent with this observation. As discussed in \S \ref{sec_results_co} above, MMS3 is driving a NW-SE outflow detected both by the SMA \cojtwo\ observations presented here and single-dish \cojtwo\ observations presented by Green et al.~(2011). The axis of this outflow may be precessing over time. \subsubsection{MMS4} MMS4 is not associated with a \emph{Spitzer} infrared source at $3.6-24$ \um, it is not associated with a \emph{Herschel} infrared source at 70 and 100 \um, and it is not associated with a SHARC-II 350 \um\ submillimeter source. The upper limits for $3.6-8$ \um\ are again taken from Guieu et al.~(2009) as described above for MMS1. For 24 \um, we take the upper limit to be the point at which the source count histogram presented in Figure 9 of Rebull et al.~(2011) turns over, which we estimate to be at a magnitude of 8.5 (corresponding to a flux density of 2.8 mJy). While there is no \emph{Herschel} 70 or 100 \um\ source or SHARC-II 350 \um\ source, there is emission at the position of MMS4 from the brighter nearby sources (MMS3 and MMS5). Thus, similar to MMS2 above, we calculate the upper limits as the flux in one beam at the position of MMS4 from other, nearby sources. With no detections of a compact, infrared source in any of the \emph{Spitzer} bands, and also no detections at $70-350$ \um, MMS4 is possibly a starless core heated only externally and thus too faint to detect at 350 \um\ above the emission from nearby, brighter sources. The 350 \um\ upper limit of 4 Jy is consistent with this statement since starless cores and cores containing very low luminosity protostars are typically less than 1 -- 2 Jy at this wavelength (Wu et al.~2007). However, recent work suggests that detections of starless cores with current interferometers are extremely rare since starless cores are not yet very centrally condensed and are thus fully resolved out (Schnee et al.~2010; Offner et al.~2012). If MMS4 is indeed a starless core, the SMA detection indicates it may be very evolved and close to the onset of star formation. Alternatively, star formation may have already begun, with the core harboring a very young, very low luminosity protostar or first hydrostatic core. Confirmation would require either the detection of a very faint infrared source below the detection limits of the data considered here, such as the 70 \um\ detection of the source Per-Bolo 58 presented by Enoch et al.~(2010), or the detection of a molecular outflow driven by this core, such as the detections of outflows from cores previously believed to be starless presented by Chen et al.~(2010), Dunham et al.~(2011), Pineda et al.~(2011), and Schnee et al.~(2012). No such outflow is detected in our \cojtwo\ observations, but this topic should be revisited with future observations providing higher sensitivity and higher spatial resolution. \subsubsection{MMS5} MMS5 is not associated with a \emph{Spitzer} infrared source at $3.6-24$ \um, and the upper limits listed in Table \ref{tab_photometry} are determined as described above. MMS5 is associated with a source detected at 70 and 100 \um\ with \emph{Herschel} and 350 \um\ with SHARC-II, and the photometry at these wavelengths is calculated with apertures and methods identical to MMS1 above. The SED of MMS5 resembles that of a Class 0 protostar too deeply embedded to be detected in the mid-infrared with \emph{Spitzer}. With no such detections we are unable to calculate $\alpha$, but calculate a value for \tbol\ (15 K) consistent with that of a Class 0 source. As noted in \S \ref{sec_results_co}, there is weak redshifted emission extending to the northwest of MMS5 and weak blueshifted emission to the southeast. These weak features may be due to an outflow driven by MMS5, but higher sensitivity \co\ observations are required to confirm this tentative outflow. \subsubsection{MMS6} Similar to MMS4, MMS6 is not associated with a \emph{Spitzer} infrared source at $3.6-24$ \um, it is not associated with a \emph{Herschel} infrared source at 70 and 100 \um, and it is not associated with a SHARC-II 350 \um\ submillimeter source. The $3.6-24$ \um\ upper limits listed in Table \ref{tab_photometry} are determined as described above. The 70, 100, and 350 \um\ upper limits are calculated as the flux in one beam at the position of MMS6 from other, nearby sources. The same discussion presented above for the evolutionary status of MMS4 also applies for MMS6; it is likely either an evolved starless core close to the onset of star formation or a very young, very low luminosity protostar or first hydrostatic core. As with MMS4, we do not detect any evidence for an outflow driven by MMS6. \subsubsection{MMS7} MMS7 is associated with the \emph{Spitzer} infrared source SST J205817.06$+$435316.1 from Guieu et al.~(2009) and Rebull et al.~(2011) and is detected at all five \emph{Spitzer} wavelengths. It is also detected at 70 \um\ with \emph{Herschel}, and the photometry presented in Table \ref{tab_photometry} is calculated with an aperture and methods identical to MMS1 above. It is not detected at either 100 \um\ with \emph{Herschel} or 350 \um\ with SHARC-II; upper limits are again calculated as the flux in one beam at the position of MMS7 from other, nearby sources. The SED of MMS7 resembles that of a young stellar object (YSO) surrounded by a circumstellar disk but no longer embedded within a dense core; the calculated values of $\alpha$ ($-$0.20; in the Flat Spectrum category) and \tbol\ (328 K; near the Class I/II boundary) confirm this assessment. \subsection{Evolutionary Status of HBC722}\label{sec_discussion_hbc722} Prior to outburst HBC722 was regarded as a Class II T Tauri star with a spectral type of K7$-$M0, a mass of $\sim 0.5-0.6$ \msun, a visual extinction of 3.4 magnitudes, an infrared spectral index of $-0.77$, and a bolometric luminosity of 0.85 \lsun\ (Cohen \& Kuhi 1979; \kospal\ et al.~2011; Miller et al.~2011). As noted in \S \ref{sec_results_continuum}, HBC722 itself is not detected in our SMA 1.3 mm continuum observations down to a $3\sigma$ upper limit of 5 mJy beam$^{-1}$. Under the same assumptions as discussed above in \S \ref{sec_results_continuum}, this corresponds to an upper limit of 0.02 \msun\ for the disk mass (lower if the disk is warmer than 30 K) and an upper limit of 3\% -- 4\% for the ratio of disk to stellar mass (again, lower if the disk is actually warmer than 30 K). In a large submillimeter survey of circumstellar disks around young stars, Andrews \& Williams (2005) report a mean disk mass of 0.005 \msun\ with a large dispersion of $\sim$0.5 dex, and a median ratio of disk to stellar mass of 0.5\%. Thus, our observations do not rule out the presence of a typical mass disk. In a typical FUor flare with an accretion rate of 10$^{-4}$ \msun\ yr$^{-1}$ and a duration of 100 yr, up to 0.01 \msun\ of mass can accrete from the disk onto the protostar. For HBC722, however, \kospal\ et al.~(2011) calculated a burst accretion rate of only 10$^{-6}$ based on their measured \lbol\ during the burst\footnote{This measurement is rather uncertain due to the variability of the outburst brightness since the initial flare in 2010 (see \S \ref{sec_hbc722}) and the fact that \kospal\ et al.~lacked photometry during the burst at $\lambda > 10$ \um, although the latter point is mitigated by the fact that the \emph{Herschel} 70 \um\ flux density of 412 mJy measured during the burst and reported by Green et al.~(2011) is generally consistent with the assumptions made by \kospal\ et al.~to extrapolate beyond 10 \um.}. Our derived upper limit for the disk mass of 0.02 \msun\ is thus consistent with providing a sufficient mass reservoir to support the observed outburst in HBC722 unless the accretion rate is several orders of magnitude higher than estimated by \kospal\ et al.~and/or the burst duration is much longer than the typical 100 yr. Even if one of these cases were true, there is possibly gas remaining in the vicinity of HBC722 that could still accrete onto the star+disk system and power the burst, as suggested by \coojtwo\ and \emph{Herschel} far-infrared continuum emission spatially coincident with HBC722. Further constraints on the amount of circumstellar mass available to accrete onto HBC722 and the likelihood of sufficient mass remaining to power further bursts beyond the current one require deeper millimeter continuum data probing to lower disk masses and higher-resolution far-infrared and submillimeter continuum data observed during the burst to better determine the burst luminosity and implied accretion rate. For the former, the high sensitivity of full-science ALMA operations will be the ideal facility despite the high declination of HBC722 ($+$44\degree) since, according to the ALMA sensitivity calculator\footnote{Available at https://almascience.nrao.edu/call-for-proposals/sensitivity-calculator}, even a short, 30 minute track with 50 antennas will improve the mass sensitivity by a factor of 100. For the latter, the upcoming 25-m submillimeter telescope CCAT will be of particular value assuming the burst is still in progress when CCAT begins science operations (currently expected in 2015 -- 2017; Radford et al.~2009; Sebring 2010). In this paper we have presented 230 GHz Submillimeter Array continuum and molecular line observations of the newly discovered FUor candidate HBC722. We summarize our main results as follows. \begin{enumerate} \item Seven 1.3 mm continuum sources are detected in the vicinity of HBC722; none are HBC722 itself. We compile infrared and submillimeter continuum photometry of each source from previous studies and conclude that three are Class 0 embedded protostars, one is a Class I embedded protostar, one is a Class I/II transition object, and two are either starless cores or very young, very low luminosity protostars or first hydrostatic cores. \item A northwest-southeast outflow is detected in the \cojtwo\ observations. This outflow is centered on and thus likely driven by MMS3, one of the Class 0 sources detected in the 1.3 mm continuum data, and its axis may be precessing. This outflow detection is consistent with a similar outflow detected in low-resolution, single-dish \cojtwo\ observations presented by Green et al.~(2011). Our higher spatial resolution confirms that HBC722 is not the driving source. \item There is no conclusive evidence that HBC722 itself is driving an outflow, although we caution that higher spatial resolution, better sensitivity to extended emission, and better determinations of the systemic velocities of the sources in the vicinity of HBC722 are needed to fully evaluate the kinematics of the \cojtwo\ gas in this region. \item The non-detection of HBC722 in the 1.3 mm continuum observations suggests an upper limit of 0.02 \msun\ for the mass of the circumstellar disk, consistent with typical T Tauri disks. This upper limit is consistent with a disk that provides sufficient mass to power the burst. Future observations are needed to further study the actual amount of circumstellar mass available to accrete onto HBC722 and the likelihood of sufficient mass remaining to power additional bursts beyond the current one. \end{enumerate} We have noted in the text several future observations that are needed in order to better disentangle the millimeter emission in this complicated environment and better determine the properties and evolutionary status of HBC722.
| 12
| 6
|
1206.4730
|
1206
|
1206.0055_arXiv.txt
|
Milliarcsecond VLBI maps of regions containing 6.7~GHz methanol maser emission have lead to the recent discovery of ring-like distributions of maser spots and the plausible hypothesis that they may be tracing circumstellar disks around forming high mass stars. We aimed to test this hypothesis by imaging these regions in the near and mid-infrared at high spatial resolution and compare the observed emission to the expected infrared morphologies as inferred from the geometries of the maser rings. In the near infrared we used the Gemini North adaptive optics system of Altair/NIRI, while in the mid-infrared we used the combination of the Gemini South instrument T-ReCS and super-resolution techniques. Resultant images had a resolution of $\sim$150\,mas in both the near-infrared and mid-infrared. We discuss the expected distribution of circumstellar material around young and massive accreting (proto)stars and what infrared emission geometries would be expected for the different maser ring orientations under the assumption that the masers are coming from within circumstellar disks. Based upon the observed infrared emission geometries for the four targets in our sample and the results of SED modeling of the massive young stellar objects associated with the maser rings, we do not find compelling evidence in support of the hypothesis that methanol masers rings reside in circumstellar disks.
|
The formation of OB stars is an important challenge to modern astrophysics as they are responsible for many of the energetic phenomena in galaxies. But their large distances, heavy obscuration and rapidity of evolution make studies of massive star formation difficult. One property that is unique to massive young stellar objects is the presence of methanol masers in their early formative phases \citep{m91,c95}. As such, they may hold clues to understanding the fundamental differences between low and high mass star formation. High angular resolution observations of the masers allow us to map the neutral gas at scales of a few 10 AUs in the vicinity of massive young stars. However, because of their relatively sparse sampling, it is still unclear where and how the masers are formed. Two competing hypotheses have arisen; one stating that methanol masers are embedded in circumstellar tori or accretion disks around the massive protostars (e.g., Norris et al. 1993), the other stating that they may generally be tracing outflows \citep{d03}. However, milliarcsecond scale very long baseline interferometry (VLBI) observations show a wide range of morphologies for the maser spot distributions at 6.7\,GHz. They form simple structures (a single maser spot), or group into spots that lie in lines or arcs, or are distributed randomly without any regularity \citep{m00,n98,p98,w98}. Linear structures of maser emission accompanied by clear velocity gradients have led to the belief that they may be in edge-on disks \citep{m00,w98}. Other scenarios where outflows or shocks are generating the linear or arc-like distributions have also been proposed \citep{d04}. We have recently completed a survey of 31 sources at 6.7\,GHz using European VLBI Network (EVN) \citep{b09}. In addition to the curved and complex morphologies observed in other samples, we have discovered for the first time nine sources (29\,\% of the sample) with ring-like maser distributions with typical sizes having major axes of 0\farcs2-0\farcs3. Though not apparent in the radio data, these ring-like structures strongly suggest the existence of a central stellar object, and prompt the obvious question: {\it Are these maser rings tracing disks around massive protostars?} Inspection of mid-infrared (MIR) data from the {\it Spitzer} IRAC maps, GLIMPSE and MIPSGAL\footnote{http://irsa.ipac.caltech.edu/}, revealed that all of our sources with ring-like morphologies coincide with unresolved MIR sources within one pixel in a GLIMPSE map (1\farcs2). Many sources also have bright near-infrared (NIR) counterparts unresolved in the 2MASS survey (Cutri et al. 2001). However, the resolutions of the {\it Spitzer} data (2\farcs3 at 8\,$\mu$m, 7$\arcsec$ at 24\,$\mu$m) and 2MASS data (2.5\arcsec) are inadequate in helping to understand the detailed environment of these maser rings. Furthermore, several of the sources as seen by {\it Spitzer} are saturated and are so bright that they cause image artifacts preventing one from knowing the true nature of any extended dust emission, if any. Therefore, we decided to obtain the highest spatial resolution NIR and MIR imaging available to explore the true nature of the maser rings. Using the Gemini 8-m telescopes, we obtained data with NIRI/Altair, an adaptive optics NIR instrument that can achieve resolutions of $\lesssim$150 mas at 2~$\mu$m. MIR observations were made with T-ReCS, employing a method which fully characterized the system point spread function accurately enough to allow the imaging data to be reliably deconvolved, achieving spatial resolutions of $\sim$150 mas at 8~$\mu$m and $\sim$250 mas at 18~$\mu$m. To better understand the nature of the methanol rings, the main aim of our observations was to resolve the circumstellar dust emission at NIR and MIR wavelengths to verify whether or not their infrared morphologies are consistent with the hypothesis that methanol maser rings are tracing circumstellar disks. Additionally, through the use of spectral energy distribution (SED) model fitting, we will learn more about the possible physical properties and source geometries of the massive young stellar objects (MYSOs) associated with the maser rings.
|
The goals of this paper were to resolve the NIR and MIR emission at the locations of four methanol maser rings to verify whether or not their IR morphologies are consistent with the hypothesis that methanol maser rings are tracing circumstellar disks, and to derive physical properties and geometries for their associated massive young stellar objects through SED model fitting. In this article we discussed the assumed distribution of circumstellar material around such young and massive accreting (proto)stars, and what infrared emission geometries would be expected for different disk/outflow orientations. For the four targets we observed, we compared the expected infrared geometries (as inferred from the properties of the maser rings) to actual high spatial resolution near-infrared and mid-infrared images. We find that the observed infrared emission geometries are not consistent with the hypothesis that the masers are residing in circumstellar disks. Using SED model fitting, we found that the emission from the infrared counterparts for all methanol masers ring distributions are indeed consistent with massive young stellar objects with masses above 8\,M$_{\sun}$. Furthermore, we find that in most cases the geometries allowed by the SED model fits corroborate the negative results from the observed infrared morphologies, casting further doubt on the hypothesis that the methanol maser rings in these four cases arise from within circumstellar disks. \\
| 12
| 6
|
1206.0055
|
1206
|
1206.3440_arXiv.txt
|
{We report the results of our numerical simulation of classical-dissipative dynamics of a charged particle subjected to a non-markovian stochastic forcing. We find that the system develops a steady-state orbital magnetic moment in the presence of a static magnetic field. Very significantly, the sign of the orbital magnetic moment turns out to be {\it paramagnetic} for our choice of parameters, varied over a wide range. This is shown specifically for the case of classical dynamics driven by a Kubo-Anderson type non-markovian noise. Natural spatial boundary condition was imposed through (1) a soft (harmonic) confining potential, and (2) a hard potential, approximating a reflecting wall. There was no noticeable qualitative difference. What appears to be crucial to the orbital magnetic effect noticed here is the non-markovian property of the driving noise chosen. Experimental realization of this effect on the laboratory scale, and its possible implications are briefly discussed. We would like to emphasize that the above steady-state classical orbital paramagnetic moment complements, rather than contradicts the Bohr-van Leeuwen (BvL) theorem on the absence of classical orbital diamagnetism in thermodynamic equilibrium.}
|
The Bohr-van Leeuwen (BvL) theorem on the absence of classical orbital diamagnetism in theromodynamic equilibrium has been a surprise of theoretical physics [1-3]. This BvL null result is statistical-mechanically exact. It is, however, counter-intuitive inasmuch as a charged particle orbiting classically under the Lorentz force exerted by an externally applied magnetic field is equivalent to an amperean current loop in the interior of the sample, and hence to a non-zero orbital magnetic moment. Moreover, the orbital moment is expected to be diamagnetic following the Lenz's law. A physically appealing resolution was advanced by Bohr [2] suggesting an exact cancellation, on the average, of the diamagnetic orbital magnetic moment in the interior of the sample by the paramagnetic moment subtended by the particle skipping the boundary in the opposite (paramagnetic) sense $-$ the edge current [3]. More recently [4], the role of the boundary was treated explicitly through a solution of the Fokker-Planck equation associated with the classical Langevin dynamics of the charged particle in a finite but unbounded space, namely the surface of a sphere (recall that, strictly speaking, a boundary has no boundary). Again, the orbital magnetic moment turned out to be zero. In our recent work [5,6] we have, therefore, re-examined the BvL null result. Our analyses strongly suggest that the vanishing of the classical orbital diamagnetism is a direct consequence of detailed-balance (the microscopic reversibility), namely that there are {\it no cycles} in thermodynamic equilibrium. Now, the detailed balance is, of course, conditioned mathematically by the second fluctuation-dissipation (II-FD) theorem of Kubo [7]. In terms of the classical Langevin equation governing the stochastic dissipative dynamics that underlies equilibrium statistical mechanics, the F-D relation requires the driving noise to be markovian. Indeed, it has been shown [8] that a markovian non-equilibrium steady state can always be transformed to an equivalent thermodynamic state in equilibrium. It is our conjecture, therefore, that a stochastic dissipative system driven by a non-markovian noise may have in general a non-equilibrium steady state with finite orbital magnetic moment $-$ without conflicting with the BvL theorem for thermodynamic equilibrium. In what follows, we have addressed this issue through a detailed numerical simulation of the relevant stochastic dissipative dynamics involved. The results of our simulation support our conjecture, namely that the orbital magnetic moment is indeed finite for the non-markovian case. There is, however, a new surprise now, namely that the sign of the magnetic moment turns out to be paramagnetic! Morever, for a classical gas of such charged particles with paramagnetic orbital moment, the inherently positive feedback may lead to an enhanced magnetic susceptibility $-$ possibly even to a spontaneous ordering of the classical orbital magnetic moments.
|
The paramagnetic sign of the induced orbital magnetic moment can have interesting consequences of considerable physical significance. After all, the paramagnetic sign of the moment inherently signifies a positive feed-back effect when we consider not just one but a system of many charged particles. Here, the mean self-field can in principle lead to a spontaneous macroscopic orbital magnetic moment. As for a possible experimental realization of such a confined system, we begin by noting that what is really essential for obtaining the classical orbital paramagnetism in a non-equilibrium {\em steady state} is the non-markovian nature of the stochastic forcing. Thus, a micron-sized sample of a semi-metal (such as bismuth) trapped in an optical tweezer, and irradiated with random laser pulses in the presence of an external magnetic field would constitute a possible candidate system. The laser impulses should impart high enough kinetic energy (high nominal temperature) so as to create a non-degenerate (classical) gas of charged particles (electrons and holes). Then, the high temperature washes out the quantum signature, i.e., the discrete quantum level-spacings in the micron-sized sample, leaving behind a classical charged particle system. (Recall that the orbital magnetic moment does not depend on the sign of the charge on the particle, e.g., be it an electron or a hole). It is then reasonable to expect the total induced orbital magnetic moment to scale up with the number of charged particles in the confined system. This mechanism may even generate spontaneously a macroscopic magnetic meanfield $-$ such as the {\em seed} field of interest in the astrophysical context.
| 12
| 6
|
1206.3440
|
1206
|
1206.3506_arXiv.txt
|
\noindent We describe a model of dark matter halo abundances and clustering which combines the two most widely used approaches to this problem: that based on peaks and the other based on excursion sets. Our approach can be thought of as addressing the cloud-in-cloud problem for peaks and/or modifying the excursion set approach so that it averages over a special subset, rather than all possible walks. In this respect, it seeks to account for correlations between steps in the walk as well as correlations between walks. We first show how the excursion set and peaks models can be written in the same formalism, and then use this correspondence to write our combined excursion set peaks model. We then give simple expressions for the mass function and bias, showing that even the linear halo bias factor is predicted to be $k$-dependent as a consequence of the nonlocality associated with the peak constraint. At large masses, our model has little or no need to rescale the variable $\delc$ from the value associated with spherical collapse, and suggests a simple explanation for why the linear halo bias factor appears to lie above that based on the peak-background split at high masses when such a rescaling is assumed. Although we have concentrated on peaks, our analysis is more generally applicable to other traditionally single-scale analyses of large-scale structure.
|
\label{intro} \citet{ps74} argued that the abundance of nonlinear virialized objects at late times (such as the present) should be sensitive to the statistics of the initial fluctuation field, and to the subsequent expansion history of the universe. This is the basis for studies which seek to use the abundance and clustering of galaxy clusters to constrain cosmological parameters. Their work has motivated the study of analytical models for the formation, and hence the abundance and spatial distribution, of halos, which can be used to provide fitting formulae when interpreting data. Following \cite{st99}, the most widely used fitting formulae are self-similar, in the sense that the predicted halo abundances can be scaled to a universal form which is independent of cosmology, redshift and power spectrum. This vastly simplifies cosmological analyses. (This universality is only expected to hold approximately, and the next generation of datasets may have sufficiently many clusters that departures from universality must be accounted for. We will have more to say about this later.) The self-similar functional form can be derived from a physically motivated model of collapse \citep*{smt01}. The number density $\der n/\der m$ of halos in the mass range $(m,m+\der m)$ is written as \be \frac{m}{\bar\rho}\frac{{\rm d}n(m)}{{\rm d}m} {\rm d}m = f(\nu)\,\der\nu, \label{excsetansatz} \ee where $\bar\rho$ is the background density and $\nu=\delta_c/\sigma$, with $\delc$ the rescaled time variable and $\sigma$ the rescaled mass variable ($\sigma^2(m) \equiv \avg{\del^2(m)}$ is the variance of the matter density field smoothed on a Lagrangian length scale corresponding to mass $m$ and linearly extrapolated to present day). Universality is manifest in the statement that $f$ depends only on $\nu$, but, unfortunately, the most naive use of this form predicts too few massive clusters. This has motivated the following ad-hoc approach: one actually fits $f(\sqrt{q}\nu)$ to the data, and determines $q$ from the fit. This semi-empirical approach has worked rather well, in the sense that $q$ appears to be approximately independent of cosmology, redshift and power spectrum, although recent simulations are beginning to show departures from universality \citep{bkk09}. Since observations will soon deliver large cluster catalogs over a range of redshifts, it is clearly desirable to have a more fundamental understanding of why $q\ne 1$, particularly because, on an object by object basis, the physical model of collapse almost never has $\delta<\delc$. I.e., $q<1$ appears to arise in the step which converts from the physics of halo formation to a statistical description of halo abundances \citep{smt01}. One of the main goals of this paper is to provide some insight into the origin of this factor. To do so, we revisit the two most common models for identifying halos from the initial fluctuation field: the peaks theory of \citet[hereafter BBKS]{bbks86}, and the excursion set approach of \citet{bcek91}. Although both make predictions which can be phrased in terms of the self-similar variable $\nu$, and both treat $\nu$ as the ratio of $\delta_c/\sigma$, the former treats the numerator of this ratio as the stochastic quantity, whereas for the latter, it is the denominator which can vary. They also differ fundamentally in their approach to the problem. Peaks theory seeks to describe the point process which describes the special positions in the initial conditions around which halos collapse. The excursion set approach aims only at a statistical description of the mass fraction in bound objects, and assumes that this can be done by consideration of all points in space -- not just the special ones around which halos form. Our analysis below shows how to merge the two descriptions. Section~\ref{mf-uncond} shows that the excursion set and peaks models can be written in the same formalism, and then describes our excursion set model for peaks, arguing that the result goes a substantial way towards explaining the origin of the factor of $a$. Section~\ref{mf-cond} extends this to describe the conditional function of excursion set peaks in constrained larger-scale environments, and from there builds a model for the large scale bias factors. This uses the recent work of \cite*[hereafter MPS]{mps12} to show that halo bias in our approach is generically expected to be $k$ dependent. It also shows that at high masses, our new expression for halo bias is qualitatively similar to that seen in simulations, again suggesting that our excursion set peaks model of the origin of $a\ne 1$ is reasonable. A final section summarizes our results, discusses them in the context of previous work on the relationship between excursion sets and peaks, and suggests ways in which our approach could be improved further.
|
We showed that, especially if one accounts for correlations between steps, the standard, spherical-collapse based excursion set prediction for halo mass fractions (equation~\ref{vfv-ms}) vastly underestimates what is measured in simulations. However, rescaling the spherical collapse motivated self-similar scaling variable $\nu\to\sqrt{0.7}\nu$ results in much better agreement (Figure~\ref{vfv-th}). We noted that this agreement should not be used to argue that halos formed from a spherical collapse -- at least, not until the reason for the adhoc rescaling of $\nu$ is understood. We then argued that the rescaling was related to a flaw in the usual formulation of the excursion set approach \citep{bcek91}, in which one replaces an average over spatial positions in one realization of the field with an ergodic average over many independent realizations of the field. Although we are not the first to have noted this problem, much previous work has attempted to rectify this by accounting for spatial correlations between walks. However, measurements in simulations showing that halos form around special positions in the initial fluctuation field \citep{smt01} suggest that it may be more productive to instead modify the ensemble over which one computes statistical averages. We then used peaks theory to illustrate this point, by showing how to incorporate the peaks constraint into the excursion set formalism. Specifically, peaks correspond to regions around which the curvature of the local density is modified (BBKS), and this, we argued, modifies the excursion set prediction from \eqn{vfv-ms} to \eqn{vfv-esp}. In fact, the fundamental role played by the curvature distribution $F(x)$ (equation~\ref{eqn-bbks-Fx}) in our analysis suggests that to build an accurate model of halo abundances, all one needs is a good model for the initial profile shapes from which halos form. For example, one might combine measurements of the density run around virialized halos with infall models to infer what the initial overdensity profiles must have been; having found them, one could use them instead of $F(x)$ in \eqn{vfv-esp} and so predict the halo mass fraction $f(\nu)$. This is in progress. Although our analysis has gone some way towards addressing the real cloud in cloud problem (correlated steps and correlated walks), there is more that can be done in this direction. This is because our analysis is fundamentally about taking `one small step' beyond that on which the object was defined; therefore, it does not correctly account for small objects which are embedded in much more massive objects (i.e., when the smoothing scales are rather different). Some of the nicest work in this direction is in a series of papers by \cite[and references therein]{m+98}; we are in the process of incorporating their work into our analysis. Our formulation of peaks in the excursion set language made it particularly easy to see how to build an excursion set model for peaks even when the question of which peaks are interesting depends on smoothing scale -- the analogue of moving barriers in the excursion set approach (equation~\ref{Nesp-Bs} and~Figure~\ref{esp-moving}). This may prove necessary if one wishes to incorporate the effects of the stochasticity associated with non-spherical collapse into the excursion set peaks predictions. The similarity in formulation also allowed a simple description of how peak abundances are modified if the large scale density field is constrained to be different in some way (equation~\ref{vfvcond-esp}). In turn, this allowed a simple generalization of earlier results on peak and halo bias to all orders (equations~\ref{bn-generic}--\ref{lamk-inverted}). In particular, we showed that excursion set peak bias is most easily understood in Fourier space, where it is $k$-dependent even at the linear level. Although we concentrated on an excursion set analysis of peaks, the MS `one-step' argument should apply to other traditionally single-scale analyses of cosmological datasets. For example, since the argument is not restricted to three dimensional fields, it can be applied to interpret the CMB temperature distribution, which is a two dimensional (nearly if not exactly Gaussian) random field. The number density and clustering of `hotspots', as a function of spot temperature, has been used as a diagnostic of the Gaussianity of this field \citep{be87, hs99}. But since some hot spots will be local maxima on larger smoothing scales as well, it is of interest to describe how the distribution of sizes (and the clustering) of regions which lie above some threshold temperature depends on the value of threshold. Clearly, the analysis presented here can be applied to that problem directly. In three dimensions, perhaps the most interesting connection and application is to the series of recent papers on the `skeleton' of the cosmic web \citep{skeleton}. This is the subject of ongoing work, where we hope to make a connection to the multi-scale analyses of \citet*{ac10}. Although essentially all the analysis in this paper used Gaussian smoothing filters (section~\ref{notation} discussed why, in the present context, they simplify the analysis substantially), we do not think they are otherwise special, so we are in the process of extending our results to include tophat smoothing filters. Since fitting functions for halo counts in simulations use tophat filtering (for the conversion between $\sigma$ and halo mass) exclusively, until our analysis does the same, a direct comparison with measurements of halo mass functions in simulations is premature. This is particularly interesting in view of the fact that the linear bias factor in our excursion set peaks model is close to the usual excursion set predictions associated with rescaled $\delc$ at small masses, but with no rescaling of $\delc$ at high masses (Figure~\ref{esp-bias}). This last is in qualititive agreement with measurements of halo bias in simulations. We believe that matching the enhanced abundance and bias at large masses (\figs{esp-vfv} and~\ref{esp-bias}), without having to rescale the parameter which is associated with the physics of halo formation, are nontrivial and encouraging successes.
| 12
| 6
|
1206.3506
|
1206
|
1206.0733_arXiv.txt
|
The Bar is the most productive region of the Small Magellanic Cloud in terms of star formation but also the least studied one. In this paper we investigate the star formation history of two fields located in the SW and in the NE portion of the Bar using two independent and well tested procedures applied to the color-magnitude diagrams of their stellar populations resolved by means of deep HST photometry. We find that the Bar experienced a negligible star formation activity in the first few Gyr, followed by a dramatic enhancement from 6 to 4 Gyr ago and a nearly constant activity since then. The two examined fields differ both in the rate of star formation and in the ratio of recent over past activity, but share the very low level of initial activity and its sudden increase around 5 Gyr ago. The striking similarity between the timing of the enhancement and the timing of the major episode in the Large Magellanic Cloud is suggestive of a close encounter triggering star formation.
|
The Small Magellanic Cloud (SMC) is a fundamental laboratory to study the evolution of dwarf irregular galaxies (dIrr's). The SMC is the closest member of this class of systems, has a current metallicity (Z$\,\simeq 0.004$ as derived from HII regions and young stars) similar to that of the majority of dIrr's and a mass \citep[between 1 and $5 \times 10^9\, M_{\odot}$, e.g. ][and references therein]{kallivayalil06} at the upper limit of their range. Moreover the SMC is a member of a triple system, a circumstance that favors studying the modulation of the star formation activities driven by interactions. In order to derive the detailed, spatially-resolved star formation history (SFH) of the SMC we have started an international long-term project to study the evolution of the SMC in space and time \citep[see][]{tosi08}. Our strategy is to achieve high photometric depth and spectroscopic resolving power over a large field of view by combining datasets from the ground and space. We are using the Hubble Space Telescope (HST), the Very Large Telescope (VLT), and the VLT Survey Telescope (VST) to observe a large sample of field stars and clusters across the SMC. These data will allow us to constrain the global SFH as well as the existence of chemical and age gradients. For the cluster analysis, we have already presented deep photometry with HST's Advanced Camera for Surveys (ACS) of seven intermediate-age and old populous clusters (\citealt{glatt08a, glatt08b, glatt09, glatt11}). In combination with our VLT data we find a complex age-metallicity relation for these clusters with a considerable spread in metallicity at any given age (see e.g. \citealt{glatt08b}). Concerning the SMC field analysis, our plan is to have Color-Magnitude diagrams (CMDs) several magnitudes fainter than the oldest main-sequence (MS) turn-off (TO) for the entire galaxy. To this purpose we have observed six fields with the ACS \citep{sabbi09}, sampling regions characterized by different stellar and gas densities in the SMC Bar, in the Wing in the direction of the LMC, and in the outskirts (see Fig. \ref{fields}). A preliminary SFH analysis of such fields has been presented in \citet{sabbi09}. We also have an ongoing Guaranteed Time Observation program at the VST \citep{ripepi06} designed to cover with deep photometry the whole SMC and the Bridge connecting it to the LMC. These CMDs will allow us to infer for the first time the SFH of the whole SMC over the entire Hubble time, covering a much larger area with considerably better image quality than \citegenitive{zaritsky2002} data. We will derive the SFHs from the CMDs using the synthetic CMD technique \citep[see e.g.][and references therein]{tolstoy09,cignoni10}. \begin{figure*} \centering \includegraphics[width=9cm]{fig1.eps} \caption{Spatial distribution of the six observed fields (red symbols) together with the observations from \citet{dolphin01} (blue symbols), \citet{mccumber05} (magenta symbol), \citet{noel07} (green symbols), \cite{chiosi07} (yellow symbols), superimposed on the DSS image of the SMC.} \label{fields} \end{figure*} SFHs of some SMC fields have been derived and presented by other authors, based on ground-based observations or HST studies of small individual regions (see Fig. \ref{fields}). \citet{harris04} derived the SFH of the SMC over $4\degr \times 4.5\degr$ to a depth of $V\la 21$ using the Magellanic Cloud Photometric Survey (MCPS) UBVI catalog \citet{zaritsky2002}. This is currently the most spatially extended study of the galaxy, but does not reach the oldest MSTO. The most comprehensive study of the old population of the SMC to date was carried out by \cite{haschke12} using RR Lyrae stars from the Optical Gravitational Lensing Experiment (OGLE-III; \citealt{udalski2008}). They find a uniform metallicity distribution across the SMC with a spread of more than 1 dex in [Fe/H]. \citet{dolphin01} analyzed the stellar content of the SMC halo, in a region close to the globular cluster NGC~121, using both HST Wide Field Planetary Camera 2 (WFPC2) and ground based data. Again with WFPC2 \citet{mccumber05} studied the stellar content of a small portion of the SMC Wing. \cite{chiosi07} derived the SFH in the vicinity of a few SMC clusters. Finally, \citet{noel07} and \cite{noel09} presented a deep ground-based study of 12 fields of the SMC, avoiding the densest regions, because of their high crowding conditions. In this paper we present the SFH of SFH1 and SFH4, the two most central fields of the six SMC regions we observed with HST/ACS \citep{sabbi09}. The apparent distances from the SMC optical center are about 24$^{\prime}$ and $1^{\circ}\,52^{\prime}$ respectively. SFH1 is located in the SW portion of the SMC Bar, where the stellar density, gas and dust contents are highest, while SFH4 is located to the NE of the SMC center, at 24$^{\prime}$ south of NGC~346, the most active star-forming region in the SMC. For a better assessment of the intrinsic theoretical uncertainties, the SFH is derived using two completely independent procedures for the application of the synthetic CMD method. We compare the two methods here and discuss the corresponding results. The other ACS fields observed by us and described by \cite{sabbi09} are being treated in the same way, and their SFH will be presented in a forthcoming paper. We briefly describe our data in Section 2. The two procedures for the SFH derivation are summarized in Section 3, together with the results of their application to SFH1 and SFH4. Similarities and differences between the resulting SFHs are discussed in Section 4, while in Section 5 we compare our findings with previous literature. Concluding remarks follow in Section 6.
|
This paper is the first of a series devoted to quantitative reconstruction of the SMC SFH from the deep HST/ACS observations presented in \cite{sabbi09}. Here we have explored the directions SFH1 and SFH4, both located in the SMC's Bar, by comparing the observational CMDs with a library of model CMDs incorporating photometric uncertainties and incompleteness as estimated by \cite{sabbi09}. To provide a robust characterization of the SFH, the choice of the best model CMD was independently conducted with two objective statistical methods, namely Cole's \citep{cole07} and Bologna's \citep{cignoni10} procedures. Our best simulated CMDs exhibit an overall good agreement with observational CMDs. The star counts along the MS and the SGB morphology are generally well reproduced, indicating that our recovered SFHs and metallicity distributions are reasonable. However, while SFH4 CMD is well fitted, there are some difficulties to reproduce the exact morphology of SFH1's CMD, especially the upper-MS spread and RC counts, which are underestimated and overestimated respectively. Concerning the resulting SFHs, a good consistency is found between the two methods in both fields. The only significant difference is the stronger rate suggested by Cole's SFH4 solution between 1.5 and 3~Gyr ago. The combination of synthetic CMDs which most resembles the observations suggests the following picture. At early times, both fields experienced a long quiescent phase characterized by low SFRs, followed by a rapid SF increase around 5-6 Gyr ago. Since then, the mode of star formation has been somewhat different in the two fields. In SFH1 the star formation was gasping and reasonably high up to today. In SFH4 it was smoother and slowly declining. To account for these differences and similarities possible explanations are: \emph{Recent burstiness:} The different level of burstiness is not surprising because these fields are separated by a distance (850~pc) larger than most of the star forming complexes discovered in the SMC (see \citealt{livanou2007}), thus allowing the recent activity to fluctuate independently; \emph{Recent systematic behavior:} The stellar density in the SFH4 region is lower than in SFH1. Hence the systematic decrease of SF activity in SFH4 may be connected with a minor amount of fuel available to support it up to today; \emph{Early quiescence and prompt rise:} Are the quiescent period and rapid SF increase 5-6 Gyr ago a global property of the Bar? Recent observations of RGB stars have revealed that older stellar components of the SMC have a velocity dispersion of about 27.5 km s$^{-1}$ \citep{harris06}, high enough to distribute the stars over a large distance (of the order of few kpc) from their birth places within few Gyr. Hence, the low early activity and the prompt rise are not peculiarities of our fields but global features of the SMC Bar. Moreover, \cite{subra2012} find no evidence for a Bar in older stars, which is consistent with our low early activity. Further support is also provided by \cite{chiosi07}, who found similar star formation trends for three other Bar fields located around the SMC clusters K~29, NGC~290, and NGC~265. From the theoretical point of view it is not clear what mechanism is responsible for the rapid rise of stellar production. Was it externally triggered by the MW or the LMC, or self-initiated? The striking similarity between the SMC and LMC SFH is suggestive of the former. Pointing in this direction are the recent calculations by \cite{diaz2011} who presented evidence that around 5.5 Gyr ago the LMC and SMC were within 160 kpc of the MW and 200 kpc of each other, therefore arguing for an independent origin of the Clouds. In this scenario the transition between quiescent and active phases could be naturally explained in terms of growing rate of mutual interactions started around 5 Gyr ago. Finally, \cite{rafe2005} noted that the spatial distribution of the younger and older clusters in the SMC is statistically different, leading to the inference that a significant accretion or merger event may have taken place around 3$-$5 Gyr ago. The study reported in this paper was the first step in a wider research activity aimed to characterize the SMC SFH through deep HST/ACS observations. A forthcoming paper will be dedicated to the analysis of the Wing and Halo fields. This will allow a comparative analysis to look for global physical characteristics in the SMC star formation process.
| 12
| 6
|
1206.0733
|
1206
|
1206.5320_arXiv.txt
|
Telescope Point Spread Function (PSF) quality is critical for realising the potential of cosmic weak lensing observations to constrain dark energy and test General Relativity. In this paper we use quantitative weak gravitational lensing measures to inform the precision of lens optical alignment, with specific reference to the Dark Energy Survey (DES). We compute optics spot diagrams and calculate the shear and flexion of the PSF as a function of position on the focal plane. For perfect optical alignment we verify the high quality of the DES optical design, finding a maximum PSF contribution to the weak lensing shear of 0.04 near the edge of the focal plane. However this can be increased by a factor of approximately three if the lenses are only just aligned within their maximum specified tolerances. We calculate the E and B-mode shear and flexion variance as a function of de-centre or tilt of each lens in turn. We find tilt accuracy to be a few times more important than de-centre, depending on the lens considered. Finally we consider the compound effect of de-centre and tilt of multiple lenses simultaneously, by sampling from a plausible range of values of each parameter. We find that the compound effect can be around twice as detrimental as when considering any one lens alone. Furthermore, this combined effect changes the conclusions about which lens is most important to align accurately. For DES, the tilt of the first two lenses is the most important.
|
Weak lensing cosmic shear has great potential to be one of the most powerful tools available to uncover the nature of dark energy \citep{detf,esoesa}. A number of planned and forthcoming surveys plan to use this probe of cosmology, including imminent surveys (KIlo-Degree Survey: KIDS, Hyper Suprime-Cam (HSC) survey\footnote{http://www.naoj.org/Projects/HSC/HSCProject.html} and the Dark Energy Survey: DES\footnote{\tt{http://www.darkenergysurvey.org}}), telescopes under construction (the Large Synoptic Survey Telescope: LSST\footnote{\tt{http://www.lsst.org}}), and future space telescopes (Euclid\footnote{\tt{http://sci.esa.int/euclid}} and WFIRST\footnote{\tt{http://wfirst.gsfc.nasa.gov}}). \begin{figure} \centering \includegraphics[width=80mm]{PSfigs/LensLayout3.eps} \caption{Diagram of the optical layout for DECam, shown in the orientation it will take when at the prime focus on La Blanco. Two light rays are coming from the primary mirror below the optical corrector. The blue light ray is an on-axis beam and the green ray is an off-axis beam. In our study, the lenses are decentred in the y direction and tilted around the z direction.} \label{fig:LensLayout} \end{figure} In the case of DES, weak lensing is one of four independent methods which will be used to determine the dark energy equation of state parameter, \textit{w}, to a precision of better than 5$\%$. The three other methods are galaxy cluster surveys, galaxy angular clustering and supernovae light curves. Details have been described in \cite{NOAO}. In weak lensing, galaxy shapes are distorted by the curvature of intervening space-time caused by matter in the Universe. As the majority of matter is dark and therefore is difficult to see by other means, lensing provides a useful tool. The gravitational lensing effect can usually be described by a slight squashing of galaxy images, called shear, and a slight bending, called flexion. The vast majority of effects are extremely small; for example an intrinsically circular galaxy would typically be sheared into an ellipse with a major to minor axis ratio of about 1.01. Weak lensing directly probes the gravitational potential along a line of sight. Unfortunately, the atmosphere and telescope imaging also distort galaxy images, and this is usually a much bigger effect than the gravitational lensing effect we are trying to measure. The most important effect of the atmosphere and telescope is well described by a convolution of the image with the point spread function (PSF). The PSF can be measured from images of point sources (stars); if the PSF is well known then it can in principle be removed, allowing a noisy estimate of the galaxy shear and flexion to be recovered. However, in practice it often leaks into the lensing measurements \citep[e.g. due to model bias or noise bias][]{voigtb10,kacprzakzrbravh12}. A key concern in developing the Dark Energy Survey programme, then, is ensuring that the Dark Energy Survey camera, DECam, has an optical system which does not add substantial image distortions. DECam has been constructed, with careful optical alignment carried out. Installation is now taking place during the first half of 2012, and the survey is due to start in Autumn 2012. The large lens size and weight make building work challenging due to the tight tolerances on the positioning. These tolerances have been designed to keep the instrumentation distortions to a minimum, so that their contributions to the weak lensing systematics is minimal, but even so, the expected ellipticity of the PSF could be larger than the lensing signal, so careful modelling of the optical distortions will be needed to recover cosmological information. In order to ensure that aspects of the alignment have been prioritised in proportion to the influence they will have on weak lensing systematics, we have engaged in the studies presented in this paper. We examine what impact errors in lens alignments have on the camera contribution to the lensing signal. The work shown here is specific to the DECam optical design, but also acts as a case study for development of future lensing-optimised optical telescopes. This work follows on from earlier work on PSF requirements from lensing by \citet{2006SPIE.6269E.100K}, who computed PSF distortions for DES for a variety of different contributions from the optics and compared the result with those from other telescopes. This paper is organised as follows. In section~\ref{sec:optics} we review the design of DECam. We discuss the shear and flexion measures used to evaluate the optical performance in section~\ref{sec:metrics}. We present our results in section~\ref{sec:results}, and discuss the implications for lensing measurements in section~\ref{sec:final}. \begin{figure*} \centering \includegraphics[width=14cm]{PSfigs/PicsOfPSFs.ps} \caption{The camera point spread function at the centre of the focal plane (top left), and at the edge of the focal plane (bottom left). Right panels show the corresponding PSFs combined with a $0.7''$ FWHM Gaussian, to describe the PSF including atmospheric seeing.} \label{fig:PSFs} \end{figure*} \begin{table} \centering \begin{tabular}{|c|c|c|c|c|} \hline & \multicolumn{2}{|c|}{Assembly Tolerances} & \multicolumn{2}{|c|}{Dynamic Tolerances} \\ & De-centre & Tilt Tolerance & De-centre & Tilt Tolerance \\ & Tolerance & on diameter & Tolerance & on diameter \\ Lens & ($\mu$m) & arcsec ($\mu$m) & ($\mu$m) & arcsec ($\mu$m) \\ \hline C1 & 100 & 10 (48) & 25 & 5.6 (27) \\ C2 & 50 & 17 (56) & 25 & 8.1 (27) \\ C3 & 100 & 20 (63) & 25 & 8.4 (27) \\ C4 & 100 & 20 (58) & 25 & 8.6 (25) \\ C5 & 200 & 40 (105) & 25 & 10 (25) \\ \hline \end{tabular} \caption{De-centre and alignment tolerances of the lenses for assembly and stability during operation, from Doel et al. 2008. } \label{tab:tolerances} \end{table} \begin{table} \centering \begin{tabular}{|c|c|c|c|c|} \hline & Centre Thickness & Edge Thickness & Diameter & Weight \\ Lens & (mm) & (mm) & (mm) & (kg) \\ \hline C1 & 112.1 & 74.34 & 980 & 172.36 \\ C2 & 51.285 & 148.11 & 690 & 87.20 \\ C3 & 75.1 & 38.27 & 652 & 42.62 \\ C4 & 101.68 & 52.48 & 604 & 49.69 \\ C5 & 54.68 & 36.19 & 542 & 24.37 \\ \hline \end{tabular} \caption{Dimensions of the DECam lenses, from Doel et al. 2008.} \label{tab:lensSize} \end{table}
|
In this paper we have explored the effect of camera optical alignments on quantities important for weak gravitational lensing, with specific reference to the Dark Energy Survey (DES). We have calculated the shear and flexion of the PSF as a function of position on the focal plane, for a realistic raytracing model of DECam including misalignments of lenses. In this work, we do not include flexure errors, or spacing tolerances which can be overcome by adjusting focus. We also do not consider flatness, position or tilt of the focal plane. However, we include lens decentres and tilts, which are known to be within tolerance for the actual DECam. For perfect optical alignment, we find a maximum PSF shear of 0.04 near the edge of the focal plane. However this can be increased by a factor of roughly three if the lenses are only just aligned within their maximum specified tolerances. In both cases, however, once PSF correction methods are employed, the resulting impact on shear estimates for galaxies in the DES weak lensing catalogue are expected to be below the required threshold for measuring dark energy parameters. We have calculated the E and B-mode shear and flexion variance as a function of de-centre or tilt of each lens in turn. We found tilt accuracy to be a few times more important than de-centre, depending on the lens considered. Finally, we have considered the combined effect of de-centre and tilt of multiple lenses simultaneously, by sampling from the permitted range of values of each parameter. In this case we find that combined effect can be around twice as detrimental as when considering any one lens alone. Furthermore, this combined effect changes the conclusions about which lens is most important to align accurately; for DES, the tilt of the first two lenses is the most important. The results of these simulations have been used to inform alignment of the DECam lenses (i.e. on which lens alignments to spend the most effort), and acts as an important confirmation that the DECam optical tolerances lead to a system which is fit for extremely accurate weak lensing measurements. The final image quality of course will be affected by other factors such as barrel sag, optical wedges, surface figure errors, focal plane misalignment and material inhomogeneity (in roughly decreasing order of significance). For example, the quality of the lens polishing introduces surface figure errors. We have assumed perfect polishing for this work. The smoothness of the final DES lenses have been examined using interferometry and photography to assess phase deviations and is well within specifications. Preliminary investigations suggest that typical atmosphere convolved PSF ellipticities are increased by up to twenty per cent in a way which varies in a non-symmetric way across the field of view. It appears to be a small but non-negligible fraction of the effect due to random misalignments of lenses within their tolerances studied in this work. Work is ongoing to incorporate all the image quality effects into the optical model along with the actual measured mis-alignments of the optics to produce a prediction of the final expected image quality.
| 12
| 6
|
1206.5320
|
1206
|
1206.7113_arXiv.txt
|
Recent observational surveys have uncovered the existence of super-luminous supernovae (SLSNe). While several possible explanations have been put forth, a consensus description for SLSNe has yet to be found. In this work we study the light curves of eight SLSNe in the context of dual-shock quark novae. We find that progenitor stars in the range of 25-35 $M_{\sun}$ provide ample energy to power each light curve. An examination into the effects of varying the physical properties of a dual-shock quark nova on light curve composition is undertaken. We conclude that the wide variety of SLSN light curve morphologies can be explained predominantly by variations in the length of time between supernova and quark nova. Our analysis shows that a singular H$\alpha$ spectral profile found in three SLSNe can be naturally described in the dual-shock quark nova scenario. Predictions of spectral signatures unique to the dual-shock quark nova are presented.
|
\label{intro} The standard astrophysical explanation for a supernova (SN) is that the radiated power is generated by energy deposited in an expanding ejecta through one of three mechanisms: the SN shock travels through the stellar envelope \citep{grassberg71}, radioactive decay of heavy elements synthesized in the explosion \citep{arnett82} or a collision with hydrogen-rich circumstellar material (CSM) \citep{chev82}. In 2011 astronomers working on the Palomar Transient Factory announced the emergence of a new class of SNe that cannot be explained by any of these means \citep{quimby11}. As described by \cite{quimby11} this new class of SNe is at least ten times brighter than a typical type Ia SN, displays spectra with little to no hydrogen, emits significant UV flux over a long period of time and has a late stage luminosity evolution that is inconsistent with radioactive decay. While this hydrogen-poor class of super-luminous SNe (SLSNe) is recent admission, the phenomenon of SLSNe as a whole has been an open question since the discovery of SN 2006gy \citep{quimby07}. Large scale supernovae surveys such as the Palomar Transient Factory (PTF) \citep{rau09, law09}, the ROTSE Supernova Verification Project (RSVP, formerly the Texas Supernova Search) \citep{quimby05} and the Catarina Real-Time Transient Survey \citep{drake09a} have uncovered approximately ten other SLSNe, some of which contain hydrogen in their spectra (SN 2006gy \citep{quimby07}, SN 2008fz \citep{drake10}) while others are hydrogen-poor (SN 2005ap \citep{quimby07}, SN 2007bi \citep{galyam09}). In order to describe these powerful explosions our understanding of the evolution of massive stars must change. \subsection{Proposed SLSN Models} \label{proposed} One method being considered to power the radiated energy of some SLSNe is a scaled up version of a CSM interaction. A dense, massive ($\sim20$M$_{\sun}$) CSM envelope must enshroud the progenitor star at the time of SN explosion in order for the SN shock to generate enough energy to power the SLSN light curve \citep{smith08,chev11,ginzburg12,kiewe12}. Building such a CSM envelope requires a mass-loss rate of $\dot{M} > 0.1 M_{\sun} $ yr$^{-1}$ over the final 10-100 years prior to SN explosion \citep{moriya12,ginzburg12}. Two possible explanations for a mass-loss rate on this order are LBV-like mass ejections \citep{smith08,kiewe12} or common envelope phase of an interacting binary system \citep{chev12}. In common envelope description, the SN is powered by the inspiral of a compact object on the core of the companion star \citep{chev12}. As well as requiring a massive CSM envelope, the conversion efficiency of kinetic energy to radiation must be much higher than that of a typical SN ($\sim 1\%$) in order to power SLSNe light curves \citep{moriya12,ginzburg12}. Alternatively increasing the kinetic energy of the SN by approximately ten-fold would have a similar effect on the energetics as increasing the conversion efficiency. The CSM model cannot explain hydrogen-poor SLSN, only type II SLSNe with strong hydrogen lines \citep{kiewe12}. An alternative description considered for SLSNe is that the radiated energy is converted from the rotational energy of a magnetar \citep{kasen10,woosley10} inside a SN envelope. For the magnetar model to power the light curve of a SLSN, large $PdV$ losses must be avoided by delaying the conversion of the magnetar's rotational energy into radiation. The delayed injection of energy into the SN envelope must be isotropically distributed across the inner edge of the SN envelope in order to energize the entire envelope and generate enough radiated energy to power a SLSN. Whether the magnetar model can power a SLSN with the expected jet-like \citep{Bucciantini09} energy deposition has yet to be studied. Although the isotropic magnetar model can provide a decent fit to the SLSN light curve, it gives no natural explanation for the observed emission lines. Pair-instability SNe (PISNe) have as well been proposed as the underlying energy mechanism for SLSNe. In this scenario an extremely massive star becomes prone to $\gamma = 4/3$ instability, triggering a SN explosion. \cite{pan12} studied the progenitor stars for PISN and found that the mass range required for a star to end its life as a PISN is $\sim300-1000 M_{\sun}$ and that no star under $1000 M_{\sun}$ in the low red-shift Universe would be susceptible to PISN. The rise time of the PISN light curve is typically much longer than that observed in the light curves of SLSNe \citep{kawabata09}. In order to model the light curve of SN 2006gy \cite{woosley07} suggested that pair-instability ejections had occurred twice and that the second ejection caught and crashed into the first ejection. \cite{woosley07} stated that in order for this collision to occur the kinetic energy of the second ejection must be artificially increased. PISN was also suggested by \cite{galyam09} as a possible explanation for SN 2007bi because the SN showed no evidence of a CSM interaction. In the supplemental material associated with the work presented by \cite{galyam09} it was noted that essentially all the available nuclear energy of the progenitor star must be converted to kinetic energy in order for the PISN to explain the light curve of SN 2007bi. Table \ref{compModelsTable} summarizes each proposed model's explanation for a variety of SLSNe characteristics. \cite{ouyed02} first suggested that a collision between material ejected through a quark nova (QN) and the preceding SN envelope could rebrighten the SN (see section 5.4 of \cite{ouyed02}). This theory was first applied in the context of SLSNe by \cite{leahy08}. In this work we provide the QN as a possible universal explanation for SLSNe. In turn we discuss what appear to be observational indications that QNe may lay at the heart of SLSNe. Observations of the SLSNe studied in this work are introduced in section \ref{obs}. Section \ref{qn} summarizes the explosion mechanism of the QN as well as the environment which leads to a SLSN. Section \ref{dsQN} examines the physics implemented in describing the interaction between ejecta of a SN and a QN. Analysis of the effects of changing physical parameters on our model light curve is undertaken in section \ref{mod}. In section \ref{results} we compare observations of eight SLSNe (SN 2005ap, SN 2006gy SN 2006tf, SN 2007bi, SN 2008es, SN 2008fz, PTF09cnd and PTF10cwr to QNe of different physical parameters. A discussion of trends found fitting the SLSNe and spectral analysis for some targets is presented in section \ref{discuss}. Finally our conclusions as well as predicted chemical signatures of our model are discussed in section \ref{conc}. \begin{table*} \caption{Comparison of proposed SLSNe models.} \label{compModelsTable} \begin{tabular}{@{}lccccc} & \vline & & Model & \\ \hline SLSNe Property & \vline & PISN & CSM & Magnetar & dsQN \\ % \hline Energy Mechanism & \vline & $\gamma = 4/3$ instability$^{\rm a}$ & Binary merger$^{\rm b}$ & Rotational energy$^{\rm c}$ & QN explosion $^{\rm d}$ \\ Radiation Mechanism & \vline & Collision of ejecta$^{\rm a}$ & CSM/SN interaction$^{\rm e}$ & Synchrotron$^{\rm c}$ & QN/SN interaction$^{\rm d}$ \\ % Progenitor mass (M$_{\odot}$) & \vline & 313$^{\rm f}$ - 1250$^{\rm f}$ & 100+$^{\rm g}$ & 8$^{\rm h}$ - 25$^{\rm i}$ & 20$^{\rm j}$ - 40$^{\rm j}$ \\ X-rays & \vline & suppressed$^{\rm k}$ & suppressed$^{\rm l}$ & not discussed & suppressed$^{\rm j}$ \\ Hydrogen in Spectra & \vline & unlikely$^{\rm m}$ & necessary$^{\rm n}$ & not discussed & likely$^{\rm j}$ \\ Cause of long-lasting broad lines & \vline & not expected$^{\rm o}$ & velocity of SN$^{\rm g}$ & not expected$^{\rm c}$ & temperature of inner shell$^{\rm j}$ \\ Late stage luminosity & \vline & radioactivity$^{\rm a}$ & opaque outer region of CSM$^{\rm p}$ & inner bubble$^{\rm c}$ & inner shell emission$^{\rm j}$ \\ \hline \end{tabular} $^{\rm a}$\cite{woosley07}, $^{\rm b}$\cite{chev12}, $^{\rm c}$\cite{kasen10}, $^{\rm d}$\cite{leahy08}, $^{\rm e}$\cite{chev82}, $^{\rm f}$\cite{pan12}, $^{\rm g}$\cite{smith10}, $^{\rm h}$\cite{weidemann77} $^{\rm i}$\cite{fryer99}, $^{\rm j}$\cite{ouyed12}, $^{\rm k}$\cite{blinnikov08}, $^{\rm l}$\cite{chev12b}, $^{\rm m}$\cite{yungelson08}, $^{\rm n}$\cite{quimby11}, $^{\rm o}$\cite{kasen12}, $^{\rm p}$\cite{chev11}, \end{table*}
|
\label{conc} We have shown that the dsQN scenario can be used to explain the light curves of all eight SLSN targets studied. In the context of dsQNe, progenitor stars ranging between 25-35 $M_{\sun}$ provide ample energy to power the large radiated energy budget of SLSNe. We found that the physical parameter with the greatest impact on dsQN light curve morphology was the time delay between SN and QN. Shorter time delay dsQN yield a peak magnitude that is higher and a faster luminosity decay rate (narrower light curve). While for longer time delays the peak magnitude is lower and the light curve is broader. A variation in time delay in the dsQN description provides an explanation for the wide variety of SLSN light curve morphology. From our analysis we found that for shorter time delay dsQN the inner shell may not be formed. The implication being that the energy that would go into forming the inner shell may instead be lost to pressure-volume work, however further study of the dynamics of the QN-SN interaction must be undertaken. We have also examined the singular H$\alpha$ spectral line profile found in three different SLSNe observations (SN 2006gy, SN 2006tf and SN 2007bi). The broad structure of the line is accounted for by thermally broadened emission from the inner shell, while the intermediate peak and blue-side absorption feature are due to contribution from the envelope. We found that the evolution of the blue-side absorption feature in the H$\alpha$ line of SN 2006gy and SN 2006tf is consistent with diffusion of the envelope. Unique to the dsQN scenario is the fact that any core collapse SN that leaves behind a massive neutron star can in turn undergo a QN explosion. This is due to the fact that the conditions of the interior of the progenitor star determine whether the neutron star could become susceptible to QN collapse. There is no correlation to the progenitor star envelope (for example whether or not hydrogen is present), thus we expect a wide variety of types of SN can become super-luminous due to re-brightening via a QN collision. A distinguishing feature predicted by the dsQN model is a unique chemical abundance caused by the spallation of the SN envelope by the QN ejecta \citep{ouyed11}. Recent observations have suggested plausible signs of the QN were found in Cas A \citep{hwang12}. As described in \cite{ouyed11} the layer of the SN envelope undergoes spallation by the QN ejecta depends on the density of the envelope at the moment of impact. For shorter time delay QN the Ni layer would be destroyed in favour of the production of sub-$^{56}$Ni elements such as; Ti, V, Cr and Mn \citep{ouyed11}. This process should also lead to mixing that would cause Fe to be found in the outer regions of a dsQN remnant \citep{ouyed11}. The formation of $^{44}$Ti at the expense of $^{56}$Ni will play an important role in the very late stage ($t\sim 1000$ days) luminosity of dsQN. For longer time delay QN the spallation process occurs in the carbon \& oxygen layer of the SN envelope which leads to a unique chemical signature of the dsQN. Spallation of the carbon \& oxygen layer would lead to an over-abundance of lithium in the dsQN remnant \citep{ouyedCEMP}. A dsQN is expected to emit two bursts of X-rays. The first X-ray emission event would occur when the shock from the original SN breaks out of the stellar envelope and the second analogously occurs for the QN shock break-out of the SN envelope. If the time delay between SN and QN is short then the X-ray bursts could in fact be overlapped leading to a broadened X-ray light curve. However if the time delay is long then there should be two distinct X-ray peaks. \footnotesize{
| 12
| 6
|
1206.7113
|
1206
|
1206.5927_arXiv.txt
|
Along with the development of formulations and numerical techniques, as well as progress in computational resources, numerical relativity (NR) is now the most viable approach for exploring phenomena accompanying strong gravitational fields, such as gravitational collapse of massive stellar core to a black hole (BH) or a neutron star (NS) and coalescence of compact-star binaries. These phenomena show a wide variety of observable signatures, including electromagnetic radiation, neutrinos, and gravitational radiation, and observations of neutrinos and gravitational radiation will provide us unique information of strong gravity and properties of dense nuclear matter otherwise cannot be obtained. Next-generation kilo-meter-size gravitational-wave detectors such as LIGO~\cite{LIGO}, VIRGO~\cite{VIRGO}, and KAGRA~\cite{LCGT} will report the first detection of gravitational waves in the next $\sim 5$ years. In addition, the above phenomena are promising candidates of the central engine of long gamma-ray bursts (LGRB) and short gamma-ray bursts (SGRB)~\cite{GRB}. All four known forces of nature are involved and play important roles in the stellar core collapse and the merger of binary compact objects: General relativistic gravity plays a crucial role in the formation of a BH and a neutron star. Neutrinos produced by weak-interaction processes govern the energy and chemical evolution of the system. The electromagnetic and strong interactions determine the thermodynamical properties, in particular equation of state (EOS) of the dense nuclear matter. Strong magnetic field, if it is present, can modify the dynamics of the matter motion. To study the dynamical phenomena in general relativity, therefore, a multi-dimensional simulation incorporating a wide variety of physics is necessary. We performed a simulation of stellar core collapse to a neutron star~\cite{Sekiguchi2010} and a black hole~\cite{Sekiguchi2011}, incorporating a finite-temperature, a self-consistent treatment of the electron capture, and neutrino cooling by a detailed leakage scheme, for the first time. Such multi-dimensional simulations had not been done in full general relativity until quite recently~\footnote{There are a number of simulations of stellar core collapse in spherical symmetry~\cite{SpheGR} in which Boltzmann's equation is solved and detailed microphysical processes are implemented.}. Ott et al.~\cite{Ott2007} (see also Dimmelmeier et al.~\cite{Dimm2007}) performed fully general relativistic simulations of stellar core collapse, employing a finite-temperature EOS derived by Shen et al.~\cite{Shen1998} (Shen-EOS) for the first time. In their calculation, however, the electron capture rate was not calculated in a self-consistent manner and neutrino cooling is not taken into account. Instead, they adopted a simple parameterized prescription proposed by Ref.~\citen{Lieb2005}: The electron fraction is assumed to be a function of density which is presumed based on a result of a single core collapse simulation with a specific initial condition. Recently, M\"uller et al.~\cite{Muller2012} performed simulations of stellar core collapse with detailed microphysics and neutrino transfer. However it is done in the framework of an approximate general relativistic gravity~\cite{xCFC}. Kuroda et al.~\cite{Kuroda2012} have made a fully general relativistic code with an approximate treatment of neutrino transfer, applying the schemes developed by the authors~\cite{Sekiguchi2010,SKSS2011}. Ott et al.~\cite{Ott2011} performed simulations of rotating stellar core collapse, employing the parameterized prescription~\cite{Lieb2005} in the collapse phase, and a ray-by-ray neutrino leakage scheme after bounce. As for the compact-star binary mergers, there have been only a few studies in general relativistic frameworks~\footnote{There are several studies of binary neutron star mergers in Newtonian frameworks in which finite-temperature EOS and weak interactions are taken into account together with neutrino cooling~\cite{Ruffert,Rosswog}.}. In the framework of an approximate general relativistic gravity (the conformal flatness approximation~\cite{CFC}), Oechslin and Janka~\cite{OJM} performed simulations of binary neutron star (BNS) mergers adopting the Shen-EOS and a finite-temperature EOS by Lattimer and Swesty~\cite{LS}, but they did not take account of weak interaction processes. Duez et al.~\cite{DuezBHNS} studied effects of EOS on the dynamics of black hole-neutron star mergers (BHNS) adopting the Shen-EOS in full general relativity. Recently, Bauswein et al.~\cite{Bauswein} performed BNS simulations adopting a wide variety of EOS in the conformal flatness approximation and investigated the dependence of gravitational wave spectra on EOS. In these works, however, the weak interaction processes are not included. In this paper, we describe our latest results of numerical-relativity simulations for the stellar core collapse to a BH~\cite{Sekiguchi2012} and the BNS merger~\cite{SKKS1,SKKS2}, which are performed incorporating both a finite-temperature EOS~\cite{Shen1998,Shen2011} and neutrino cooling~\cite{Sekiguchi2010}. For reviews on other topics, namely, simulations of stellar core collapse to a neutron star, of BHNS merger, and of BH-BH binary merger, the reader may refer to Refs.~\citen{Ott,Kotake,LLR}, Refs.~\citen{ShibataTaniguchi,Duez} and Refs.~\citen{Pretorius,Centrella}, respectively. The paper is organized as follows. In \S~\ref{Sec_BasicEq}, we first briefly summarize basic equations, input microphysics, and numerical setup. The results of simulations of the stellar core collapse and BNS merger are described in \S~ \ref{Sec_ResultSCC} and \S~\ref{Sec_ResultBNS}, respectively. Section~\ref{Sec_Summary} is devoted to a summary. Throughout this paper, $\hbar$, $k_{B}$, $c$, and $G$ denote the Planck's constant, the Boltzmann's constant, the speed of light, and the gravitational constant, respectively. In appendices, details of the microphysics adopted in our latest implementation are summarized for the purpose of completeness. We adopt the geometrical unit $c=G=1$ in \S~\ref{Sec_3+1} and \S~\ref{Sec_Leakage}.
|
\label{Sec_Summary} We have described our latest results of numerical-relativity simulations of rotating stellar core collapses to a BH and BNS mergers, performed incorporating a finite-temperature EOS and neutrino cooling effects. The following is the summary of our latest findings and prospects for the near future. \subsection{Stellar core collapse} We presented our latest results of axisymmetric simulations of very massive stellar core collapsing to a system composed of a rotating BH and surrounding disk/torus in full general relativity. The simulation were performed taking into account of the microphysical processes and the neutrino cooling. Because progenitor models of LGRBs suggested in the literatures~\cite{Fryer07} propose a possibility that they may have an entropy higher than that of ordinary supernova cores, we employed a model of a presupernova core with a high entropy of $s/k_B = 4$ calculated by Umeda and Nomoto~\cite{UN2008} together with hypothetical two rotational profiles (UN100-rigid and UN100-diff). As in the collapse of ordinary supernova cores, the gravitational collapse sets in due to the photo-dissociation of heavy nuclei and the electron capture. The collapsing core eventually experiences a core bounce, forming a shock wave. Then a HMNS, which is supported by the centrifugal force and the thermal pressure is, formed. The neutrino luminosity in this HMNS phase is larger than than that in the ordinary supernova as $\gtrsim 5\times 10^{53}$ ergs/s. The HMNS eventually collapses to a BH irrespective of the initial rotational profiles. However, the dynamics and the geometry of the final outcome depend strongly on the degree of the initial rotation. For the model UN100-rigid, the shock wave formed at the core bounce is deformed to be a torus-like shape. Then the infalling materials are accumulated in the central region after they pass through the oblique shock formed at the tours-shaped outer region of the HMNS. As a result, the thermal energy is efficiently stored at the surface of the HMNS due to the dissipation of the kinetic energy of the accumulated materials, driving outflows. After the collapse of the HMNS, a torus is formed around the BH. We found that the torus shows a time variability. The total neutrino luminosity emitted from the torus around the BH amounts to $L_{\nu, {\rm tot}}\sim 10^{51}$--$10^{52}$ ergs/s, which lasts for a long duration of $\gtrsim 1$ s. Associated with the time variability of the BH-torus system, the neutrino luminosities also show a violent time variability. Such a long-term high luminosity with the time variation may be related to the time variability that LGRBs show. For the model UN100-diff, by contrast, a geometrically thin disk is formed around the BH and the BH-disk system shows essentially no time variability. Remarkably, the above differences in the dynamics and the outcome stem from a small difference in the initial rotational profile. We also calculated the characteristic gravitational-wave strain $h_{\rm char}$ for UN100-rigid. The effective amplitude is as large as $\sim 10^{-20}$ at $f \sim 1$ kHz for a hypothetical event occurred at a distance of 10 kpc. This shows a possibility that we may observe multi-messenger information, namely, gravitational waves, neutrinos, and electromagnetic radiation from such a nearby event and may obtain a clue to understand the stellar core collapse and the central engine of LGRBs. \subsection{Binary neutron star merger} We showed that for a stiff, purely nucleonic EOS, a HMNS is the canonical outcome and a BH is not promptly formed after the onset of the merger as long as the total mass of the system is smaller than $3.2M_{\odot}$. The primary reason is that the thermal pressure plays an important role for sustaining the HMNS. We further showed that the lifetime of the formed HMNS with mass $\lesssim 3M_{\odot}$ would be much longer than its dynamical time scale, i.e., $\gg 10$~ms, and will be determined by the time scale of the subsequent neutrino cooling. The neutrino luminosity in the early evolution phase of the HMNS was shown to be high as $\sim 3$--$10\times 10^{53}$~ergs/s. The effective amplitude of gravitational waves averaged over the random source direction and orbital plane inclination is $h_{\rm eff}=4$--$6 \times 10^{-22}$ at $f_{\rm peak}=2.1$--2.5~kHz for a hypothetical source distance of $D=100$~Mpc. If the BNS merger happens at a relatively short source distance $\sim 20$~Mpc or is located in an optimistic direction with $D \sim 50$~Mpc, such gravitational waves may be detected by advanced gravitational-wave detectors with ${\rm S/N}=5$, and the HMNS formation will be confirmed. For an EOS in which effects of hyperons are taken into account, the EOS becomes softer than the purely nucleonic EOS. With this EOS, a BH is often formed in a short time scale after the onset of the merger, although a HMNS could be a transient outcome with a short lifetime $\lesssim 10$~ms. Because the EOS becomes soft during the evolution of the HMNS, the compactness significantly changes in a short time scale in this EOS. This is well reflected in gravitational waveforms and their spectra. Specifically, the characteristic frequency changes with time. This effect reduces the amplitude at a peak frequency of gravitational waves in the Fourier space, and make a feature unfavorable for the detection of gravitational waves. Roughly speaking, the allowed distance for the detection of gravitational waves from the HMNS is by a factor of 2 smaller than that in the nucleonic EOS for the same mass of BNS. \subsection{Future prospects} \subsubsection{Massive stellar core collapse} As mentioned in \S~\ref{Sec_ResultSCC}, we did not take into account effects of the neutrino heating in the numerical simulations to date. Recently, we have developed~\cite{SKSS2011} a formulation for numerical simulations of general relativistic radiation transfer based on Thorne's moment formalism~\cite{Thorne1981}. Based on this formalism, we have already performed general relativistic radiation magnetohydrodynamics (GRRMHD) simulations~\cite{SS2012} for the evolution of a system composed of a BH and a surrounding torus with a simplified treatment of microphysics, as a step toward a more physical modelling. Furthermore, we have succeeded in implementing a code which can solve the neutrino transfer with a detailed microphysics (in preparation). Using this code, we plan to perform simulations of the stellar core collapse to explore a supernova explosion mechanism and the formation of a BH in full general relativity. Throughout the study of the massive stellar collapse in this article, we assume the axial symmetry in the simulations. However, this assumption would be invalid if non-axisymmetric instabilities set in. For example, in Ref.~\citen{Kiuchi2011}, it has been shown that the BH-torus systems could exhibit the so-called Papaloizou-Pringle instability~\cite{PP}. Also, for a rapidly rotating HMNS, non-axisymmetric instabilities such as a bar-mode instability could set in~\cite{SS05,Rampp98,Ott3D,Simon,LLR}. Once these instabilities turn on, the torus/HMNS may deform to be a highly non-axisymmetric structure. This will enhance the angular momentum transport in the torus and HMNS, and the evolution processes of these systems may be modified. We plan to perform a three dimensional simulation to explore if the non-axisymmetric instabilities set in and play an important role for the collapsar models. \subsubsection{Binary neutron star merger} To date, we performed the BNS merger simulations for the case of the equal-mass binaries. As a straightforward extension of the previous studies, we plan to perform simulations for unequal-mass binaries, for which the merger dynamics, gravitational waveforms, and the mass of the disk will be modified. We also plan to perform merger simulations of BHNS binaries. It has been reported (see Ref.~\citen{Kyutoku2011} and references therein) that a massive disk of $M_{\rm disk}\gtrsim 0.1M_{\odot}$ can be formed for stiff EOSs even when the BH rotates moderately ($a_{\rm BH} \gtrsim 0.5 $). Such a system is a promising candidate of the central engine of SGRBs. Only simulations with detailed microphysics will enable a quantitative study for this merger hypothesis of SGRB. In the study of the BNS merger simulation, we totally ignore the effect of magnetic fields. Recent studies on magnetized BNS merger showed that the magnetic fields in the merger and post-merger phases have an impact on the dynamics of the torus formed around the BH~\cite{LR}. This is because the angular velocity inside the torus has a steep gradient. Thus, the magnetic field is subject to the amplification via magnetic winding and/or magneto-rotational instability~\cite{Balbus}. This condition holds in the HMNS because it has strong and rapid differential rotation as discussed in \S~4.2. We plan to incorporate magneto-hydrodynamics in our code and explore its impact on the evolution of HMNSs.
| 12
| 6
|
1206.5927
|
|
1206
|
1206.4173.txt
|
Previous hydrodynamical (HD) and magnetohydrodynamical (MHD) numerical simulations of hot accretion flows have indicated that the inflow (gas with inward radial velocity) accretion rate decreases with decreasing radius. Two models have been proposed to explain this result. In the adiabatic inflow-outflow solution (ADIOS), the inward decrease of accretion rate is because of the loss of gas in the outflow. In the alternative convection-dominated accretion flow (CDAF) model, the accretion flow is thought to be convectively unstable; the gas is then assumed to be locked in convective eddies, which results in the inward decrease of the accretion rate. In the present paper we investigate the nature of inward decrease of accretion rate using HD and MHD simulations. We calculate various properties of inflow and outflow, including the mass flux, radial and rotational velocities, temperature, and the Bernoulli parameter ($Be$). Systematic and significant differences between inflow and outflow are found. For example, for HD flows, the temperature of outflow is significantly higher than inflow; while for MHD flows, the specific angular momentum of outflow is nearly Keplerian, which is significantly higher than inflow. These results suggest that the inflow and outflow are not dominated by convective turbulence, but they are systematic inward and outward motion. We have also analyzed the convective stability of MHD accretion flow and found that they are convectively stable. These results indicate that the inward decrease of inflow rate is because of the mass loss in outflow. The different properties of inflow and outflow also suggest that the mechanisms of producing outflow in HD and MHD flows are buoyancy associated with the convection and centrifugal force associated with the angular momentum transport mediated by the magnetic field, respectively. The latter mechanism is similar to the Blandford \& Payne mechanism but no large-scale open magnetic field is required; it is kind of ``micro-Blandford \& Payne'' mechanism. We also study the effect of initial conditions in the simulations. We find that the value of $Be$, whose sign determines whether the outflow can escape to infinity or not, is mainly determined by the value of $Be$ of the initial condition. We discuss some possible observational applications, including the Fermi bubble observed in the Galaxy center and winds widely observed in active galactic nuclei and black hole X-ray binaries.
|
There are two series of black hole accretion solution, depending on the temperature of the accretion flow. The cold series include the standard thin disk and the slim disk, bounded by the Eddington accretion rate $\dot{M}_{\rm Edd}\equiv 10 L_{\rm Edd}/c^2$ (Sunkura \& Sunyaev 1973; Abramowicz et al. 1988). For the hot series, when the accretion rate is below $\sim \alpha^2 \dot{M}_{\rm Edd}$, where $\alpha$ is the viscous parameter, it is the advection-dominated accretion flow (ADAF; Narayan \& Yi 1994; 1995; Abramowicz et al. 1995; see reviews by Narayan, Mahadevan \& Quataert 1998, Kato, Fukue \& Mineshige 1998, and Yuan \& Narayan 2013). The main application of ADAFs is on the dim black hole sources, including the supermassive black hole in our Galactic center, low-luminosity AGNs, quiescent and hard states of black hole X-ray binaries (see reviews by Narayan 2005; Yuan 2007; Narayan \& McClintock 2008; Ho 2008). Above $\alpha^2\dot{M}_{\rm Edd}$, up to even close to $\dot{M}_{\rm Edd}$, the solution is described by the luminous hot accretion flow (LHAF; Yuan 2001; 2003). LHAFs correspond to higher accretion rates and radiative efficiency therefore the emitted luminosity is higher. This model has been applied to explain the origin of hard X-ray emission in relatively luminous sources such as luminous hard state of black hole X-ray binaries and some AGNs (e.g., Yuan \& Zdziarski 2004; Yuan et al. 2007). In the early analytical studies of ADAFs, it was {\em assumed} that the mass accretion rate, or more precisely the inflow rate (inflow and outflow rates are defined as the mass flux of gas with a negative and positive radial velocity; refer to eqs. [\ref{inflowrate}] \& [\ref{outflowrate}] in the present paper for their definitions), is a constant of radius. Correspondingly, the density follows a power-law distribution: $\rho (r) \propto r^{-3/2}$. Later many numerical simulations have been performed to study the multi-dimensional dynamics of the hot accretion flow, including both hydrodynamical (HD) and magneto-hydrodynamical (MHD) ones (e.g., Stone, Pringle \& Begelman 1999, hereafter SPB99; Igumenshchev \& Abramowicz 1999, 2000; Stone \& Pringle 2001; Hawley, Balbus \& Stone 2001; Machida, Matsumoto \& Mineshige 2001; Hawley \& Balbus 2002; Igumenshchev, Narayan \& Abramowicz 2003; Pen, Matzener \& Wong 2003; De Villiers, Hawley \& Krolik 2003; De Villiers et al. 2005; Yuan \& Bu 2010; Pang et al. 2011; McKinney, Tchekhovskoy \& Blandford 2012; Narayan et al. 2012; Li, Ostriker \& Sunyaev 2012; Yuan, Wu \& Bu 2012, hereafter Paper I). One of the most important findings of these simulations is that the inflow accretion rate decreases with decreasing radius. The radial profile of the inflow rate can be described by a power-law function of radius, $\dot{M}(r)\propto r^{s}$ with $s\sim 0.5-1$. Correspondingly, the radial profile of density becomes flatter, $\rho(r)\propto r^{-p}$ with $p\approx 1.5-s\sim 1-0.5$. Such a result not only has theoretical significance, but also is of great importance to observations. This is because it determines the emitted spectrum of the accretion flow; and also the strength of ``mechanical AGNs feedback'', if such a profile is caused by mass outflow as we will illustrate in the present paper. Because of its importance, we first need to critically examine the reliability of this result. This is because, due to technical difficulties, all the above-mentioned numerical simulations have a rather small radial dynamical range, which usually spans two orders of magnitude. In this case, the results may suffer from the boundary conditions thus are suspectable. To overcome this problem, in Paper I we presented a ``two-zone'' approach, successfully extended the dynamical range to four orders of magnitude. We found that in this case, the profiles of accretion rate and density are almost remain unchanged compared to previous works. In addition, we combined previous numerical simulation works on hot accretion flows, including both HD and MHD ones, and found that the profiles of accretion rate and density in all these works are quite similar. Such a consistency is in agreement with the prediction of a recent work by Begelman (2012), although slopes are different. What is more exciting is that these theoretical results have been confirmed by the observations, e.g., to Sgr A* and NGC~3115 (Yuan, Quataert \& Narayan 2003; Wong et al. 2011; see Paper I for details). The next question is then: what is the physical reason for the decrease of the accretion rate? Two main models have been proposed to answer this question. The first one is the adiabatic inflow-outflow solution (ADIOS; Blandford \& Begelman 1999; 2004; Begelman 2012; see also Becker, Subramanian \& Kazanas 2001; Xue \& Wang 2005; and Jiao \& Wu 2011 for related works.). In this model there is inflowing and outflowing zones with almost equal but opposite mass fluxes. The net accretion rate is orders of magnitude smaller. The inward decrease of mass accretion rate is because of the mass loss in the outflow launched at every radius. In the early versions of the model (Blandford \& Begelman 1999; 2004), the origin of the outflow was assumed to be because of the positive sign of the Bernoulli parameter of the accretion flow. But in the latest version of the ADIOS model the mechanism of producing outflow is not specified but leave open (Begelman 2012). The second model is the convection-dominated accretion flow (CDAF; Narayan, Igumenshchev, \& Abramowicz 2000; Quataert \& Gruzinov 2000). This model is based on the assumption that the hot accretion flow, both hydrodynamical (HD) and magnetohydrodynamical (MHD), is convectively unstable. In this model, when $\alpha$ is small, the inward angular momentum transport by convection and outward transport by viscous stress almost cancel each other. A convective envelope solution was then constructed which can reproduce the simulated flat density profile. In this scenario, the inward decrease of mass accretion rate is because that with accretion, more and more fluid is locked in convective eddies operating circular motion. The first aim of the present work is to investigate which one of the above two scenarios is correct. For this purpose, we run some HD and MHD simulations, and calculate the respective properties of inflow and outflow, including their radial and rotational velocities, temperature, and Bernoulli parameter. If the CDAF scenario is correct, i.e, the motion of the flow is dominated by convective turbulence, and inflow and outflow rates are simply due to turbulent fluctuation, we should expect that the properties of inflow and outflow are similar. As we will see, however, we find that the properties of inflow and outflow are systematically and significantly different. The second question we want to address is the convective stability of MHD accretion flow. The HD hot accretion flow was predicted to be convectively unstable (Narayan \& Yi 1994). Physically this is because the entropy increases with decreasing radius, which arises because of the entropy production by viscous dissipation and little entropy loss in the radiation. Such a prediction has been well confirmed by HD simulations (SPB99; Igumenshchev \& Abramowicz 1999, 2000). CDAF model further assumes that an MHD accretion flow is also convectively unstable and applies to an MHD flow as well. However, such an applicability was questioned from the beginning by some authors, because they think that the dynamics of MHD flow is controlled by magnetorotational instability (MRI; Balbus \& Hawley 1991, 1998) rather than convection (Stone \& Pringle 2001; Balbus \& Hawley 2002). Narayan et al. (2002) did a linear MHD stability analysis and argued that if the flow is unstable the long-wavelength modes of the instability are intrinsically ``convection'' so the CDAF model should be applicable to MHD accretion flows; moreover, the convective stability can be judged by the HD H{\o}iland criteria. In the present work we use the simulation data to analyze the convective stability of MHD accretion flows. We find that they are convectively stable. Based on this result, combined with the fact that HD and MHD accretion flows have almost identical inward decrease of accretion rate, we conclude that the decrease of accretion rate for both HD and MHD flows is not because of the convection, but systematic outflow. Then two immediate questions arise. The question in the theoretical aspect is: what is the physical origin of these outflow? In the aspect of observational applications, the most important question that concerns us is the main properties of outflow such as their terminal velocity and kinetic power. We investigate these two questions again by systematically comparing the properties of inflow and outflow. One potential complication of the numerical simulation is the effect of initial condition adopted in the simulation. Many simulations choose a rotation torus as the initial condition; in some other works the gas is injected into the computation domain, i.e., ``injection-type'' initial condition. The effect of the initial condition according to our knowledge was never systematically studied. Our idea of the potential importance of initial condition is stimulated by the one-dimensional steady global solution of ADAFs. Since the accretion equations are a set of {\em differential} equations, and since the differential terms in ADAFs (i.e, advection) are very important, we expect that the outer boundary condition should be important in determining the global dynamics of ADAFs. In other words, for given parameters, the solutions of ADAFs are not unique, but comprise a series of solution corresponding to different outer boundary conditions. This expectation was fully confirmed by the detailed calculations of Yuan (1999). Following this spirit, the initial condition should also play some role in multidimensional numerical simulations, i.e., the final solution should preserve some memory to the initial condition, although we don't know to what extent before the detailed calculations are done. We expect that while the main properties of accretion flow should remain, some properties of outflow may depend on the initial conditions. One particular example is the Bernoulli parameter of the accretion flow. This parameter is the sum of kinetic energy, enthalpy, and potential energy (eq. [{\ref{bernoulli}]). As we will describe later, its value will determine whether the outflow can escape to infinity and what is the terminal velocity of the outflow. In the self-similar solution of ADAFs, it was found that $Be>0$ (Narayan \& Yi 1994). However, as later pointed out by several authors the positive sign of $Be$ is only a result of self-similar solution. In the global solution, its sign should depend on the the outer boundary condition (Nakamura 1998; Yuan 1999; Abramowicz, Lasota \& Igumenshchev 2000). This was confirmed by the numerical calculations of Yuan (1999). We assign the systematic study of the effects of initial condition and boundary condition in numerical simulation of accretion flow in a separate paper (Bu, Yuan \& Wu 2012, in preparation). In this paper, we consider three different initial conditions in our HD simulations, but only one initial condition for the MHD simulation. The paper is structured as follows. The four models, one MHD and three HD ones with different initial conditions, will be described in \S2. The systematic comparisons of properties between inflow and outflow are presented in \S3. In \S4, we compare our simulation results with previous ones. We find that some puzzling results in previous works can be understood as due to different initial conditions. In \S\ref{adioscdaf}, we analyze the convective stability of MHD accretion flow, and argue that the ADIOS scenario is favored by our simulations. The origin of outflow in both HD and MHD accretion flows are discussed in \S\ref{outfloworigin}. In \S\ref{observation}, we discuss the observational application of our study. The last section (\S\ref{summary}) devotes to a summary.
|
\label{summary} Numerical simulations of hot accretion flow, both HD and MHD, have revealed that the mass accretion rate decreases with decreasing radius. Consequently the radial density profile of the accretion flow becomes much flatter compared to the self-similar solution (Narayan \& Yi 1994) which is based on assumption of a radius-independent mass accretion rate. Denoting the profiles of accretion rate and density as $\dot{M}(r)\propto r^s$ and $\rho(r)\propto r^{-p}$, the HD simulation in Paper I, which has the largest radial dynamical range so far (from $r_s$ to over $10^4r_s$), gives $s\sim 0.4-0.75$ and $p\sim 0.65-0.85$, respectively. As a comparison, the self-similar solution of ADAFs gives $s=0$ and $p=1.5$. Such a flatter density profile have obtained strong observational support in Sgr A* and NGC~3115 (see Paper I for details). Two different models have been proposed to explain such a result, namely ADIOS and CDAF. In this paper we investigate the nature of the inward decrease of accretion rate and other related questions. For this aim, we have performed a series of HD and MHD numerical simulations, comparing the various properties of inflow (the gas with a negative radial velocity) and outflow (the gas with a positive radial velocity) such as radial and rotational velocities, temperature, and Bernoulli parameter. We found systematic and significant differences (\S3). Such differences are hard to understand if the inflow and outflow are simply the appearance of turbulent fluctuation; but strongly suggest that the motion of the accretion flow is dominated by systematic inward and outward fluxes of mass. We have also analyzed the convective stability of MHD flows. We found that they are stable (\S5). This indicates that CDAF model at least can't be applied to MHD flows. Based on these results, together with other arguments, we conclude that the inward decrease of accretion rate is because of mass loss via outflow (\S\ref{adioscdaf}). An immediate question is then the origin of outflow. This is discussed in \S\ref{outfloworigin}. The detailed comparison of properties between inflow and outflow again presents important information to answer this question. In the HD case, we found that the temperature of outflow is systematically higher than that of inflow (Fig. \ref{Fig:temperature}). This suggests that the outflow is driven by the buoyancy which arises because of the convective instability of the HD accretion flow. In the case of MHD flow, we found that the specific angular momentum of outflow is very high, close to the Keplerian angular momentum at the equatorial plane; while the angular momentum of inflow is much lower (Fig. \ref{Fig:rotationvelocity}). This suggests that it is the centrifugal force that drives the production of outflow. Magnetic field in the flow efficiently transfers the angular momentum between fluid elements. Whenever one element gets enough angular momentum, it will turn around and be thrown outward. This mechanism is similar to the Blandford \& Payne (1982) mechanism, except that no large-scale magnetic field is required here. We therefore call it ``micro'' Blandford \& Payne (MBP) mechanism (Fig. \ref{Fig:mbp}). The properties of outflow is of great interest especially because of their potential important role in AGNs feedback. We have calculated the mass flux, terminal radial velocity, momentum flux and kinetic power of outflow which should be useful in the comparison with observations of AGNs winds, and in the study of AGNs feedback (\S3). One of the most important properties is the Bernoulli parameter of outflow since its sign determines whether the outflow can escape to infinity. We have run three HD models with different initial conditions. We found that while many properties of the accretion flow are not dependent of the initial condition, the sign of $Be$ of outflow is crucially determined by the value of $Be$ in the initial condition (Fig. \ref{Fig:radialbernoulli}). If $Be$ is large in the initial condition, $Be$ of outflow inclines to be positive. This is natural since the total energy should be conserved in the simulation. Unfortunately, it is uncertain what kind of initial condition is more realistic, and the answer may depend on circumstances. We think that the value of $Be$ of the gas is likely to be positive at least in some cases such as when the radiative energy loss of accretion flow is small. Even in the case that $Be<0$, these outflow can still escape out of the outer boundary of the accretion flow, as shown by our simulations; thus they can still interact with the ISM and play a similar role to those outflow with $Be>0$. We have also discussed the possible applications of the outflow from hot accretion flow in explaining observations (\S\ref{observation}). These include the formation ``Fermi bubble'' in the Galactic center, and the origin of observed winds from both AGNs and black hole X-ray binaries. These winds have been detected from sources with a variety of luminosities. Detailed analysis to some sources published in literature have shown that it is very hard for the winds to be produced by radiation or thermal driving mechanisms. Our simple estimation suggests that it is promising to explain their origin by our ``MBP'' mechanism.
| 12
| 6
|
1206.4173
|
1206
|
1206.7041_arXiv.txt
|
Hydrogen recombination lines are one of the major diagnostics of \hii\ region physical properties and kinematics. In the near future, the Expanded Very Large Array (EVLA) and the Atacama Large Millimeter Array (ALMA) will allow observers to study recombination lines in the radio and sub-mm regime in unprecedented detail. In this paper, we study the properties of recombination lines, in particular at ALMA wavelengths. We find that such lines will lie in almost every wideband ALMA setup and that the line emission will be equally detectable in all bands. Furthermore, we present our implementation of hydrogen recombination lines in the adaptive-mesh radiative transfer code RADMC-3D. We particularly emphasize the importance of non-LTE (local thermodynamical equilibrium) modeling since non-LTE effects can drastically affect the line shapes and produce asymmetric line profiles from radially symmetric \hii\ regions. We demonstrate how these non-LTE effects can be used as a probe of systematic motions (infall \& outflow) in the gas. We use RADMC-3D to produce synthetic observations of model \hii\ regions and study the necessary conditions for observing such asymmetric line profiles with ALMA and EVLA.
|
Massive stars influence their cosmic environment through powerful winds, radiation feedback and supernova explosions. Ionizing radiation from massive stars produces pronounced \hii\ regions around them. Observations of \hii\ regions, while still in the ultracompact phase (with diameters less than $0.1$~pc), provide important insight into massive star formation \citep{habis79,churchwell02,zinnyork07}. Observations of hydrogen recombination lines at radio and sub-mm frequencies (RRLs) are routinely used by observers to infer densities, temperatures and velocity structures inside \hii\ regions \citep{gorsor02}. While the microphysics of recombination lines was reasonably well understood relatively early \citep{dupreegold70}, the \hii\ regions themselves had to be modeled either on scales larger than the scale of gravitational collapse using numerical radiation-hydrodynamics \citep[see reviews by][]{yorke86,maclow07,klessenetal11} or on smaller scales with simple spherically \citep{brownetal78,keto02,keto03} or non-spherically \citep{keto07} symmetric models. Only recently, collapse simulations of massive star formation with ionization feedback have left the constriction of two dimensions \citep{richyork97} and facilitated fully three-dimensional dynamical simulations of \hii\ region expansion during massive star formation \citep{petersetal10a,petersetal10b,petersetal10c,petersetal11a}. These collapse simulations require adaptive-mesh simulations with 10 refinement levels and more. To post-process such data, dedicated radiative transfer tools must be developed. To create synthetic recombination line observations of the simulated \hii\ regions, we have implemented recombination lines in RADMC-3D. RADMC-3D is a radiative transfer code that can handle both continuum \citep[e.g.][]{petersetal10b} and line \citep[e.g.][]{shettyetal11} radiation on arbitrary octree meshes. It has an interface for PARAMESH \citep{macneiceetal00}, the grid library of the adaptive-mesh code FLASH \citep{fryxell00}, which was used for these simulations. We have used RADMC-3D previously to model free-free and dust continuum emission from the simulated \hii\ regions \citep{petersetal10b}. To the best of the authors' knowledge, the only other codes that can model recombination line observations are MOLLIE \citep{keto90}, which we have used in previous work to post-process simulation data \citep{petersetal10a} and for radiative transfer modeling \citep{longmoreetal11}, and an innominate code that \citet{mapietal93} used for recombination line models of MWC~349. The advantages of RADMC-3D are its direct compatibility with many major simulation codes and its modularity. Another important feature of RADMC-3D is the possibility to create user-defined setups. Thus, it can be used by observers interested in creating simple analytic models of the regions they are observing. In the present paper, we describe the implementation of hydrogen recombination lines in RADMC-3D and demonstrate its capability with synthetic observations of several \hii\ region models. The development of such a tool is particularly timely because of the commissioning of the Expanded Very Large Array (EVLA) and the Atacama Large Millimeter Array (ALMA). These facilities are set to open new frontiers in terms of sensitivity, angular resolution, dynamic range and image fidelity in the cm to sub-mm wavelength regimes. Tools that can help to interpret such new observational data are clearly highly desirable. The purpose of the present paper is twofold. First, we present our implementation of hydrogen recombination lines in RADMC-3D and describe a suite of tests of this implementation. The user-defined analytical setups are ideal for code validation before applying the method to the much more complex simulation data. Second, we use these analytical models to simulate synthetic EVLA and ALMA observations. We particularly discuss the appearance of line asymmetries expected for ALMA observations. The paper is organized as follows. Section~\ref{sec:physrecomb} briefly summarizes the physics of hydrogen recombination lines. In Section~\ref{sec:impl}, we describe the implementation of recombination lines in RADMC-3D. In Section~\ref{sec:rrl_properties}, we discuss the general properties of RRL transitions at ALMA frequencies. We then focus on the observation that the RRL profiles can be asymmetric (Section~\ref{sec:asymm_profiles}) and investigate the conditions under which we can expect to observe such asymmetries (Section~\ref{sec:preditc_conditions}). We conclude our paper with a summary of our results in Section~\ref{sec:summary}. Additional information on line profile asymmetries can be found in Appendix~\ref{sec:symm}. We present our code tests in Appendix~\ref{sec:tests}.
|
\label{sec:summary} We have presented a discussion of the capabilities for ALMA to observe hydrogen recombination lines and of their physical properties. An important difference between ALMA lines and lines at cm-wavelengths is that the level populations of the former lines are farther away from LTE conditions, which gives rise to the formation of asymmetric line profiles. We have investigated necessary conditions to observe such profiles with ALMA and presented synthetic ALMA and EVLA observations of simple model \hii\ regions. We find that the asymmetric line profiles can provide useful information about systematic gas motion, such as infall or outflow, within the \hii\ region that was previously unaccessible. Radiative transfer modeling, e.g. with the RADMC-3D implementation of recombination lines presented here, will be necessary to distinguish non-LTE effects from geometric effects due to asymmetries in the \hii\ region and to extract the desired information from the observations. Interested readers can obtain the code upon request from the first author.
| 12
| 6
|
1206.7041
|
1206
|
1206.5044_arXiv.txt
|
The polar motion is a phenomenon in the rotational motion of the Earth, appearing as the motion of the point on the Earth's surface at which the rotational axis pierces the surface. The motion is roughly circular but with changing radius. This phenomenon occurs in the free rotation of a rigid Earth as well. When the Earth is regarded as a rigid oblate spheroid, the locus is a complete circle. The period of the circular motion in the rigid case is $A/(C - A)$ times the period of the Earth's rotation, $C$ and $A$ being the moments of inertia of the Earth around its symmetrical axis and around an axis on the equator of the symmetrical axis, respectively. We suppose the approximation that the Earth is a spheroid throughout the present study. For the value of $A/(C - A)$ in the actual Earth, the period is about 304 sidereal days. The motion of the rotational axis in a rigid Earth is called the Euler motion and its existence and the behavior can be proved theoretically. In the actual Earth, the Euler motion suffers a considerable deformation. The motion is now not with a simple shape and with a single period but a compound of the various elliptical motions with different periods, shapes and sizes, all varying with time. The largest ellipse among them stems from the Euler motion in the rigid case and is called the free oscillation (Munk and McDonald, 1960). The period of of the free oscillation is about 435 days, considerably longer than that in the rigid Earth. The free oscillation is also called the Chandler wobble after the name of the discoverer of the period. The prolongation of the period in the free oscillation in the actual Earth is attributed to the elasticity of the Earth. It was first explained by Newcomb (1892). Explanations are also seen in Munk and McDonald (1960) as well as in Kubo (1991). As it is important to take the prolongation of the period into account in the argument in Section 2, a detailed review of it is given there. Besides the Chandler wobble and the many other periodical components in the polar motion is also found a secular drift of its center. All the motions other than the free oscillation are considered to be reflecting the variation of the mass distribution in the whole Earth including the ocean and the atmosphere. Therefore, the investigation of the polar motion has held a very important position for understanding the dynamical state of the Earth. For that reason, numerous analyses have been published concerning the polar motion, especially for its periodic parts. Among them, some studies aim only to analyze the pattern of the polar motion itself such as the variations in its amplitude and phase, e.g., Kimura (1917), Guinot (1972), H$\ddot {\textrm{o}}$pfner (2002). Some other studies intend to explain the variation in the polar motion by geophysical causes such as the atmosphere (Aoyama and Naito, 2001), the ocean (Dickman, 1993; Gross, 2000), geomagnetic jerks (Gibert, 2008) and so on. The polar motion, however, is considered to reflect the state of the Earth not so directly as the variation of the figure axis. The change in the inner state of the Earth first brings about a variation in the figure axis, which then affects the polar motion. Therefore it is the variation of the figure axis on the Earth's surface but not the polar motion that geophysical events should be linked to. The present study intends to obtain the position and the motion of the figure axis relative to the Earth's surface. It is performed by the method of rigid dynamics, regarding the elasticity and anelasticity of the Earth as perturbations to the rotation of a rigid Earth. Primarily, a main theme of the rotation of the Earth in rigid dynamics is to obtain the respective motions of the rotational axis and the figure axis. The solutions to the rotation of the rigid Earth as well as of the elastic Earth have been published numerously so far, among which a typical example for the former is Kinoshita (1977) and one for the latter would be Kubo (1991). It is not necessary at present, however, to solve the respective motions of both axes. We only have to determine the position of the figure axis at any moment on the Earth's surface and we can attain this purpose only if the relation between the rotational and the figure axes is known because the observed polar motion is the position of the rotational axis on the Earth's surface. Both the figure axis and the rotational axis are definite physical entities and the relation between them can be determined without any ambiguity. The derivation of the fundamental equations to give the relation between both axes is carried out in Section 2 based on the result in Kubo (1991), and some preliminary studies using the equations are presented in Sections 3 and 4. In Sections 5 and 6, the actual polar motion of the Earth is exermined. First in Section 5 are overviewed the feature of the polar motion published by IERS as well as the spectrum distribution for the components contained in it, with a result that all the conclusions reached are not different largely from those preceding. In Section 6, on the other hand, is presented a quite new result on the position and the motion of the figure axis which are obtained from the data of the polar motion. Finally, in Section 7, a simple model of the mass transfer in the Earth is introduced to examine if the variation of the figure axis obtained above is reasonable or not.
|
The variation of the figure axis reflects the physical state in the Earth more directly than the polar motion since the change inside the Earth first brings about a variation of the figure axis and then it affects the polar motion. In the present study, the equatons to give the relation between the figure axis and the rotational axis in the Earth with both elasticity and anelasticity have been derived. Since the polar motion is the motion of the rotational axis on the Earth's surface, the obtained relation makes it possible to determine the position of the figure axis in the Earth-fixed frame at any moment from the data of the polar motion. The position and the motion of the figure axis obtained from the actual data of the polar motion exhibits various characteristics. First, the averaged position of the polar motion are not coincident with that of the figure axis exactly but is dislocated by about $0.01''$, while the differnce between the secular motions of both is not so definite. The periodic variation of the figure axis has obvious and comparatively stable components of annual and semi-annual periods but no component with the period of the free oscillation. On the whole, the annual and the semi-annual components draw a shape resembling a thin ellipse with the semi-major axis of about $0.02''$. While the figure axis draws a locus like this as the average for a long interval, the shape changes from year to year and, other than that, there exist many minor variations with shorter periods. Those minor variations are considered to be significant but not due to the errors in the observations or in the process of the analysis, and therefore they are necessary to be studied more in detail. Finally, a simple model with the seasonal flow of the matter in the Earth has been introduced to explain the variation of the figure axis obtained above. Applying the same model also to the variation of the rotational speed, it is confirmed that the variation estimated is accordant with the one which is so far known, thus suggesting that the variations of the figure axis and of the rotational speed are brought about by the same cause as far as the seasonal variation is concerned.
| 12
| 6
|
1206.5044
|
|
1206
|
1206.2722_arXiv.txt
|
We describe the overall characteristics and the performance of an optical CCD camera system, Camera for QUasars in EArly uNiverse (CQUEAN), which is being used at the 2.1 m Otto Struve Telescope of the McDonald Observatory since 2010 August. CQUEAN was developed for follow-up imaging observations of red sources such as high redshift quasar candidates ($\mathrm{z} \gtrsim 5$), Gamma Ray Bursts, brown dwarfs, and young stellar objects. For efficient observations of the red objects, CQUEAN has a science camera with a deep depletion CCD chip which boasts a higher quantum efficiency at $0.7 - 1.1 \, \mu m$ than conventional CCD chips. The camera was developed in a short time scale ($\sim$one year), and has been working reliably. By employing an auto-guiding system and a focal reducer to enhance the field of view on the classical Cassegrain focus, we achieve a stable guiding in 20 minute exposures, an imaging quality with FWHM $\geq 0.6\arcsec$ over the whole field ($4.8\arcmin \times 4.8\arcmin$), and a limiting magnitude of $z = 23.4$ AB mag at 5-$\sigma$ with one hour total integration time.
|
Recent emphasis on the distant and/or obscured target % observation requires near infrared (NIR) sensitive imaging devices. Although the conventional CCD chips, especially the back illuminated, thinned CCD chips, are very sensitive over a wide wavelength range ($0.4 - 0.8 \, \mu m$), most of them are not sensitive in NIR bands, with the quantum efficiency (QE) of order of a few percent at $\sim 1 \, \mu m$ \citep[e.g.,][]{im10a}. Detectors made of low band-gap semiconductor materials such as HgCdTe and InSb offer high sensitivity in NIR, but the NIR arrays are expensive in cost, and have strict export restrictions, making them less accessible to many astronomical imaging applications. Considering the recent high demand for NIR imaging, it is thus important to develop a cost-effective imaging camera. To facilitate imaging out to $1.1 \, \mu m$ at a reasonable cost and within a fast development schedule, the Center for the Exploration of the Origin of the Universe (CEOU) has developed a CCD camera system, Camera for QUasars in EArly uNiverse (CQUEAN). It is a camera system designed primarily for the observation of high redshift quasar candidates. The design goal was to make a camera sensitive at $0.7 - 1.1 \, \mu m$ (QE $>$ 20 \% at $1 \, \mu m$) with a moderate field of view (at least a few arcmin across). The sensitivity in the red is governed by the need for the detection of the Lyman break of objects at $\mathrm{z} \gtrsim 5$. The requirement for the field of view is to allow photometry based on the reference stars available from large area surveys such as the Sloan Digital Sky Survey (SDSS) \citep{yor00} and the Large Area Survey of UKIRT Infrared Deep Sky Survey (UKIDSS) \citep{law07}. The optical performance requirement was to obtain % the seeing limited point spread function (PSF) with FWHM better than $1\arcsec$. Based on these requirements, we adopted a CCD camera with a deep depletion chip, a focal reducer, and an auto-guiding system. The development of CQUEAN started in the summer of 2009, and it was installed on the Cassegrain focus of the 2.1 m Otto Struve telescope at the McDonald Observatory, Texas, USA, on 2010 August 12. Since its commissioning run, CQUEAN has been used for various scientific projects. In this paper, we describe the overall characteristics of CQUEAN, and its performance based on the first year of the scientific operation. Section 2 describes the overview of the instruments including the science % camera, its filters, and the parameters of CQUEAN on the 2.1 m telescope. In Section 3, we describe various characteristics of the camera and images obtained with CQUEAN. We report the performance of CQUEAN in Section 4, and briefly describe the scientific result obtained with CQUEAN in Section 5. Finally, summary and conclusion are presented in Section 6.
|
We developed CQUEAN, an optical CCD camera system which has an enhanced sensitivity at $\sim 1 \,\mu m$ compared to the conventional CCD camera. CQUEAN consists of a science camera, seven filters, a focal reducer to enlarge its field of view, and a guide camera system. CQUEAN also includes its in-house GUI software for controlling the instrument. It is attached at the Cassegrain focus of the 2.1 m telescope at McDonald Observatory, USA. The analysis of the CQUEAN data reveals that the camera is stable and it meets the requirements set by the scientific purpose. % With the focal reducer of our own design, the camera produces images of the seeing as good as $0.6\arcsec$ without any fringe pattern in longer wavelength bands, although coma aberration is present for $g$ and $r$-band images as expected. We estimated the sensitivity of the camera to be $z = 23.4$ AB mag. at one hour integration. In addition, the characteristics of bias, dark, and flat images of the science camera are examined, and the observation and data reduction strategy is discussed to obtain better results under these system characteristics. We also find that CQUEAN can obtain images of exposures up to 1200 second with its auto-guider system. With the enhanced sensitivity at NIR together with a field of view of $4.8\arcmin \times 4.8 \arcmin$, CQUEAN has been serving as a useful instrument for the observation of red, astronomical objects such as high redshift quasar candidates, afterglows of GRBs, and supernovae.
| 12
| 6
|
1206.2722
|
1206
|
1206.6930_arXiv.txt
|
Known populations of QSOs appear to fall short of producing the ionizing flux required for re-ionizing the universe. The alternative, galaxies as sources of ionizing photons, suffers from the problem that known types of galaxies are almost completely opaque to ionizing photons. For reionization to happen, either large numbers of (largely undiscovered) sources are required, or the known populations of galaxies need to have had a much larger escape fraction for ionizing radiation in the past. We discuss recent discoveries of faint $z\sim 3$ Lyman $\alpha$ emitters with asymmetric, extended Ly$\alpha$ emission regions, which apparently are related to interacting galaxies. The unusually shaped Ly$\alpha$ line profiles and the underlying stellar populations of these objects suggest the presence of damaged gaseous halos, infall of gas, tidal or stripped stellar features and young populations of hot stars that would all be conducive to the release of ionizing radiation into intergalactic space. As galaxy interactions and mergers increase with redshift, these effects can only become more important at earlier times, and so these interacting $z\sim3$ objects may be late, lower redshift analogues of the sources of reionization.
|
The known population of QSOs, thought to be the main source of ionizing photons throughout most of the history of the universe, falls short of maintaining the ionization of the intergalactic medium at redshifts beyond $z\sim 4$ (e.g., \cite{rau97}), whereas the universe is known to be fully ionized back to at least $z\sim 6$ (e.g., \cite{bec07}). Ionizing flux escaping from galaxies is often invoked to make up for the missing photons. The highest redshift galaxy surveys are finding increasing numbers of $z>7$ galaxies. However, to reach the number density of photons necessary for re-ionization may require either the existence of large numbers of (still) invisible sources (e.g., Trenti, this conference), or rather large escape fractions for those types of galaxies currently observable (e.g., Finkelstein, this conference). The challenge posed by these conditions is exacerbated by a paucity of credible observations of any galaxies emitting a significant amount of ionizing radiation (see, e.g., \cite{van10} for a recent discussion). This may suggest that yet other types of sources may be responsible for reionization ("hidden" AGN, mini-halos), or that the galaxies at those lower redshifts currently accessible to observation for some reason are no longer leaking ionizing photons. With ordinary galaxies failing to ionize the universe, it is instructive to briefly consider possible conditions that may facilitate the release of ionizing radiation. These conditions generally fall into the following three categories: \begin{enumerate} \item enhanced production of ionizing photons, e.g., by hotter, perhaps metal poor stars, or production of harder photons that experience a lower optical depth. \item creation of holes in the otherwise optically thick galactic disks or halos, through which ionizing radiation may escape, possibly through interactions that damage a gaseous halo, galactic winds, or through ionization by an AGN. \item the formation in, or transport of the stellar sources of ionizing radiation into less optically thick environments. Examples may be any kind of intra-halo star formation, e.g., in globular clusters or tidal tails. \end{enumerate} In the absence of the difficult, direct detections of leaking ionizing continuum, the prime observational diagnostic of the above features would be the Ly$\alpha$ emission from neutral hydrogen. A Ly$\alpha$ emission line from a gaseous halo that is leaking ionizing radiation may be spatially asymmetric, and possibly offset from the underlying sources of the ionizing continuum. Density and velocity disturbances of the halo gas, caused by galaxy interactions, could produce such asymmetries, as could multiple sources of ionizing radiation moving about in a joint gaseous halo. The presence of the latter may also be evident from broad band detections of multiple stellar continuum objects, perhaps in the form of disturbed stellar populations, e.g., tidal features, and young stellar populations, the formation of which may have been triggered by interactions.
|
We have discussed some astrophysical conditions that could turn galaxies into sources of ionizing radiation during the epoch of reionization, and we described two recently discovered Ly$\alpha$ emitting galaxies that satisfy several of these conditions, including an apparently leaky gaseous halo, and possible extragalactic, intra-halo star-formation from a metal poor reservoir of gas (apparently in the form of tidal tails or turbulent wakes). The connection of these features to galaxy interactions suggest that, as the merger rate increases with redshift, such objects will be much more common at the epoch of reionization. The extended $z\sim 3$ emitters discussed here may just be the lower redshift remnants of this population. In contrast, the symmetric, compact Ly$\alpha$ emitters or Lyman break galaxies represent the bulk of the $z\sim 3$ star-forming galaxies. In the absence of recent interactions, these objects, with their more spatially concentrated star-formation, more mature stellar populations, and their ionizing radiation from a dominant central source being trapped by an inviolate gaseous halo, may be less likely to contribute to the ionization of the universe. We conclude that interactions between galaxies may be the missing ingredient that turns high redshift galaxies into the anticipated sources of ionizing radiation. In neither one of the two peculiar Ly$\alpha$ emitter cases can we exclude the presence of an AGN, which could provide alternative explanations for both the high Ly$\alpha$ widths observed, and the escape of ionizing photons. However, taking into account that there is already another QSO in the same (small) survey volume, the assumption, that all three objects could reflect photonionization by internal QSOs, would raise the number density for QSOs at $z\sim3$ to $\sim 1.5\times10^{-3}h_{70}^3$Mpc$^{-3}$. Such a high value in itself may have interesting consequences for the budget of ionizing photons. \begin{theacknowledgments} I would like to thank my collaborators George Becker, Martin Haehnelt, Jean-Rene Gauthier, Swara Ravindranath, and Wal Sargent, and acknowledge useful discussions with Sebastiano Cantalupo, Bob Carswell, Hsiao-Wen Chen, Jeff Cooke, Li-Zhi Fang, Pat McCarthy, Masami Ouchi, and Francois Schweizer. I further thank the staff of the Las Campanas and the Keck Observatories for their help with the observations, the IoA in Cambridge and the Raymond and Beverley Sackler Distinguished Visitor program for hospitality and support, and the NSF for funding through grant AST-1108815. \end{theacknowledgments}
| 12
| 6
|
1206.6930
|
1206
|
1206.4104_arXiv.txt
|
We present a detailed analysis of high resolution H~I observations of the Magellanic spiral galaxies NGC 4618 and NGC 4625. While the H~I disk of NGC 4625 is remarkably quiescent with a nearly uniform velocity dispersion and no evidence of H~I holes, there is a dynamic interplay between star formation and the distribution of neutral hydrogen in NGC 4618. We calculate the critical density for widespread star formation in each galaxy and find that star formation proceeds even where the surface density of the atomic gas is well below the critical density necessary for global star formation. There are strong spatial correlations in NGC 4618 between UV emission, 1.4 GHz radio continuum emission, and peaks in the H~I column density. Despite the apparent overlap of the outer disks of the two galaxies, we find that they are kinematically distinct, indicating that NGC 4618 and NGC 4625 are not interacting. The structure of NGC 4618 and, in particular, the nature of its outer ring, are highly suggestive of an interaction, but the timing and nature of such an interaction remain unclear.
|
One of the more stunning revelations about the Local Group in the past few years has been the observations of the proper motions of the Magellanic Clouds showing that they are moving at significantly higher velocities than previously thought (e.g. Kallivayalil et al. 2006, Piatek et al. 2008, Besla et al. 2010). One major implication of these results is that the Magellanic Clouds may very well be on their first passage through the Local Group, contradicting the long-held view that they are long-standing companions to the Milky Way. From another point of view, this argument makes sense. The Magellanic Clouds, and the Large Magellanic Cloud in particular, are unlike any other companions of the Milky Way, M31, or the vast majority of other spiral galaxies (e.g. James \& Ivory 2011). In addition, over the past decade or so we have seen a proliferation of statistically significant samples of galaxies that show that objects sharing the basic morphology of Magellanic spirals are common in both the local Universe and at intermediate redshift (Garland et al. 2004, Ryan-Weber et al. 2003, Sheth et al. 2008). Part of the motivation for the study described here is that the dynamics, structure, and star formation history of the LMC have long been interpreted in the context of its proximity to both the Milky Way and the SMC. The fact that there are so many galaxies that share the LMC's structure but not its proximity to other massive galaxies, and the possibility that the LMC has only recently entered the Local Group drives us to examine the properties of the larger population of Magellanic spirals in detail sufficient to meaningfully and properly understand how the LMC fits into the larger context of late-type barred galaxies. The earliest comprehensive look at Magellanic spirals was carried out by de Vaucouleurs \& Freeman (1972) who noted the structural similarity between the LMC and a number of other nearby galaxies. Much of the subsequent work was aimed at understanding the origin of the lopsided structure that characterized Magellanic type galaxies. Interactions were primarily believed to be responsible for the asymmetric properties of this class of galaxy; a scenario which was well fit by the LMC itself. Theoretical work and some simulations could also accurately account for the apparent lopsidedness of Magellanic spirals, but not their frequency. Baldwin et al. (1980) originally suggested that differential precession of the disk in an axisymmetric potential would give rise to the asymmetry characteristic of SBm galaxies. Their proposal, however, predicted that the winding problem would result in the asymmetry dispersing over 5 Gyr, too short a timescale to account for the ubiquity of lopsidedness. Walker, Mihos, \& Hernquist (1996) and Zaritsky \& Rix (1997) proposed that minor mergers could result in lopsided distributions. Compelling evidence for the role of interactions in accounting for lopsided morphology comes from a statistical survey based on examination of POSS and UK Schmidt plates that found 71/75 Magellanic spirals with well-defined one armed morphologies appear to have a physical neighbor (Odewahn 1994). Wilcots \& Prescott (2004), however, showed that only 4 of 13 galaxies from the Odewahn (1994) sample were actually interacting and Wilcots, Lehman \& Miller (1996) showed that even those that were interacting did so weakly. These studies weakened the argument that asymmetry is generally caused by interactions, and bolstered a model put forth by Levine and collaborators (Levine \& Sparke 1998, Noordermeer, Sparke, \& Levine 2001) that lopsidedness is stable once the disk is displaced from the dynamical center of the halo and is allowed to rotate around the center. One example of an instance in which an interaction may be linked to the characteristics of a Magellanic spiral is the NGC 4618-4625 pair of galaxies. NGC 4618 and NGC 4625 have both been classified as Magellanic-type spirals (Odewahn 1994), though only NGC 4618 exhibits the classic optical morphology of a barred galaxy in this class (de Vaucouleurs \& Freeman 1972). Tully (1988) gives a distance of 7.3 Mpc to NGC 4618 and 8.2 Mpc NGC 4625, but their apparently overlapping H~I distributions lends support to the notion that interactions play a role in generating the characteristic morphology (Bush \& Wilcots 2004). In addition, recent {\it GALEX} observations reveal NGC 4625 to have an extended UV disk as well as multiple branching, ragged spiral arms. NGC 4625 has been noted to have one of the largest UV to optical disk (as measured by $D_{25}$) ratios of 4 (Gil de Paz et al. 2005) and an even larger H~I to optical disk ratio of 9.8 (Bush \& Wilcots 2004). It is remarkably symmetric and Bush \& Wilcots (2004) noted that it seems unaffected by its apparent interaction with NGC 4618. From its morphology alone, NGC 4618 appears to be a prototypical Magellanic spiral, with an obvious stellar bar and a lopsided one-armed spiral structure. Both galaxies are host to on-going star formation; in NGC 4618 most of the H~I regions lie near the central bar, while in NGC 4625 the star formation is nearly uniformly distributed across its disk. The goals of this work are to re-examine the issue of whether these two galaxies are indeed interacting and what effect such an interaction may be having on both NGC 4618 and NGC 4625, particularly with regard to star formation. We also seek to better understand the symbiotic relationship between the neutral atomic interstellar medium and star formation in Magellanic spirals. In this paper we present a detailed analysis of high-resolution observations of the neutral gas content of NGC 4618 and NGC 4625, complementing the earlier work of Bush \& Wilcots (2004) (hereafter, BW04). We describe our observations in $\S$2 and present the H~I distribution and overall kinematics in $\S$3.1. We show the tilted ring models of rotation curves for each galaxy in $\S$3.2. We begin our analysis in $\S$4 by looking at the distribution of massive stars and the structure and kinematics of the neutral gas in each galaxy. $\S$4.1 identifies the H~I holes in the center of NGC 4618 and $\S$4.2 looks into correlations with tracers of massive star formation. $\S$4.3 calculates star formation thresholds and the corresponding measured column densities. The velocity dispersion of each H~I disk is addressed in $\S$4.4. We discuss the possibility of an ongoing interaction between NGC 4618 and NGC 4625 in $\S$5 and finish our analysis in $\S$6 where we look into the nature of the H~I ring around NGC 4618. Our conclusions are stated in $\S$7. \footnotetext{\footnotesize The National Radio Astronomy Observatory is a facility of the National Science Foundation operated under cooperative agreement by Associated Universities, Inc. This work also made use of observations made with the NASA Galaxy Evolution Explorer. {\it GALEX} is operated for NASA by the California Institute of Technology under NASA contract NAS5-98034.}
|
We presented a detailed analysis of high resolution H~I observations of NGC 4618 and NGC 4625. In combination with tracers of recent star formation these high resolution H~I data allow us to carry out an investigation of the relationship between massive star formation and the distribution and kinematics of the atomic interstellar medium in both galaxies. While the surface density of the atomic gas remains well below the nominal critical density needed for global star formation (e.g. Kennicutt 1989), we do find a spatial correlation between tracers of current star formation and relative peaks in the H~I column density. In NGC 4625 star formation traces faint spiral structure that is also seen in the H~I column density. In NGC 4618 star formation is found in much of the disk. We also noted a number of H~I holes in NGC 4618, but also show that the holes are not expanding, and establish anti-correlation between the location of the H~I holes and peaks in UV emission for the galaxy. As with a number of other studies we conclude that the major holes are most likely the result of supernovae. However, we reach this conclusion without an understanding of the molecular content of the holes, but believe that the holes are more likely to contain hot (10$^{6}$ K) gas rather than cold gas. The 1.4 GHz radio continuum emission stemming from the center of NGC 4618 suggests that the emission stems from multiple supernovae, strengthening the argument that the holes were created through a series of powerful explosions. We investigate the environments in which we observe active star formation in both galaxies and measure H~I column densities in the area of the brightest star formation corresponding to the areas of brightest UV emission. The H~I line profiles associated with the gas believed to be the bridge connecting the two galaxies actually has two well-defined and separate kinematic components. This observation makes a strong case against an on-going interaction between the NGC 4618 and NGC 4625 galaxies, in conjunction with the lack of perturbations to either velocity field or the H~I velocity dispersion in NGC 4625. The persistence of the H~I ring around NGC 4618 allows for us to apply a rough interaction timescale of 450 Myr, assuming that the ring has successfully completed one complete rotation around the center of the galaxy. Investigating the area that surrounds the NGC 4618/4625 system, we find that NGC 4490 is projected to be relatively close to the pair, yet serves as an unlikely candidate for having previously interacted with NGC 4618. Thus, the origins of the likely stable H~I ring around NGC 4618 are still not well understood.
| 12
| 6
|
1206.4104
|
1206
|
1206.6101_arXiv.txt
|
By implementing widely-used equations of state (EOS) from Lattimer \& Swesty (LS) and H. Shen et al. (SHEN) in core-collapse supernova simulations, we explore possible impacts of these EOS on the post-bounce dynamics prior to the onset of neutrino-driven explosions. Our spherically symmetric (1D) and axially symmetric (2D) models are based on neutrino radiation hydrodynamics including spectral transport, which is solved by the isotropic diffusion source approximation. We confirm that in 1D simulations neutrino-driven explosions cannot be obtained for any of the employed EOS. Impacts of the EOS on the post-bounce hydrodynamics are more clearly visible in 2D simulations. In 2D models of a 15 $M_{\odot}$ progenitor using the LS EOS, the stalled bounce shock expands to increasingly larger radii, which is not the case using the SHEN EOS. Keeping in mind that the omission of the energy drain by heavy-lepton neutrinos in the present scheme could facilitate explosions, we find that 2D models of an 11.2 $M_{\odot}$ progenitor produce neutrino-driven explosions for all the EOS under investigation. Models using the LS EOS are slightly more energetic compared to those with the SHEN EOS. The more efficient neutrino heating in the LS models coincides with a higher electron antineutrino luminosity and a larger mass that is enclosed within the gain region. The models based on the LS EOS also show a more vigorous and aspherical downflow of accreting matter to the surface of the protoneutron star (PNS). The accretion pattern is essential for the production and strength of outgoing pressure waves, that can push in turn the shock to larger radii and provide more favorable conditions for the explosion. Based on our models we investigate several diagnostic indicators of the explosion that have been suggested in the literature, e.g., the amplitude of the standing-accretion shock instability mode, the mass weighted average entropy in the gain region, the PNS radius, the antesonic condition, the ratio of advection and heating timescales, the neutrino heating efficiency, and the growth parameter of convection.
|
\label{sec:intro} Core-collapse supernova explosions are triggered by the gravitational energy released during the transition from a stellar core to a protoneutron star (PNS). The thermodynamic conditions obtained at the center of the stellar core are temperatures on the order of tens of MeV and densities on the order of normal nuclear matter density ($3\times10^{14}$~g~cm$^{-3}$). There are only few equations of state (EOS) available for these conditions. Several studies by \citet{taka82} and \citet{baro85} were pioneering, investigating the impact of the EOS on the explosion dynamics, albeit in a phenomenological approach (see collective references in \citealt{beth90}). The most commonly used nuclear EOS in recent supernova simulations are the EOS from \citet{latt91} (hereafter LS), based on the incompressible liquid-drop model including surface effects, and from \citet{shen98} (SHEN). The latter is based on relativistic mean field (RMF) theory and the Thomas-Fermi approximation. Recently, a new nuclear RMF EOS has been published in \citet{hemp10}, which in addition can be obtained for several nuclear parametrizations and provides detailed information about the chemical composition (see also \citealt{gshen,furusawa}). Thus, extensive studies have been performed to investigate the impact of the nuclear EOS on supernova simulations. But most of these simulations are limited to spherical symmetry. \citet{thom03} investigated the effects of the incompressibility parameter on the evolution of temperature and neutrino luminosities during the postbounce evolution based on the three different incompressibilities of the LS EOS ($K=180$, 220, 375~MeV). Only a small impact on the dynamics was found. \citet{sumi06,sumi07} performed long-term simulations of a 40 $M_\odot$ progenitor star from \citet{woos95} comparing LS ($K=180$~MeV) and SHEN, focusing on the evolution until black-hole formation and the emitted neutrino signal. \citet{fisc09} included in addition progenitor stars from different stellar evolution groups in the mass range of 40--50 $M_\odot$ and studied the neutrino signal and collapse to a black hole in a systematic way. Their results were confirmed and extended by \citet{ocon11} applying a simplified treatment of neutrino transport. In \citet{hemp12}, the authors applied their new EOS and systematically explored the hydrodynamic evolution and emission of neutrinos for different nuclear parameters within the class of RMF EOS. Moreover, \citet{sage09} and \citet{fisc11} explored the appearance of quark matter via a first order phase transition in supernova simulations during the early post-bounce phase, which triggers the explosion even in spherically symmetric models. Furthermore, EOS studies based on multi-dimensional supernova models have become available recently. \citet{kota04b} and \citet{sche10} performed two-dimensional (2D) and three-dimensional (3D) simulations, respectively, and investigated the initial post-bounce phase and gravitational wave emission focusing on effects of rotation and convection with simplified neutrino treatments. They found that the gravitational wave signal depends strongly on the EOS (e.g., \citealt{kota11} for a recent review). In addition, \citet{mare09b} investigated the EOS dependence of high frequency variations of the neutrino luminosity and gravitational wave emission by performing 2D simulations that include detailed neutrino-radiative transfer. Several 2D studies obtain neutrino-driven explosions of massive iron-core progenitors \citep[see e.g.][]{mare09b,brue09,suwa10,muel12b}, aided by the standing accretion-shock instability (SASI) and neutrino-driven convection. These simulations were evolved until the expanding explosion shock left the central core ($\sim$1000~km) between about 400--600~ms after bounce (depending on the progenitor and details of input physics). Most multi-dimensional supernova studies used the LS EOS with the incompressibility of $K=180$~MeV, which is a very soft EOS \citep{bura06,suwa10}. In addition, \cite{mare09} performed 2D simulation using the stiffer EOS from \citet{hill84}, and found that the softer EOS results in a more optimistic situation for a possible onset of an explosion. However, their simulation with the stiffer EOS had been performed for a shorter post-bounce time than the softer EOS. Hence, a final conclusion cannot be drawn.\footnote{Employing a stiffer EOS with $K=263$ MeV based on the Hartree-Fock approximation, Marek et al. found no explosions for the same progenitor model, whereas they indeed obtained an explosion for Shen EOS that is even stiffer with $K=281$ MeV \citep[see, e.g.,][]{jank12}.} Even further, they employed only two EOS, each of which has a different incompressibility and symmetry energy. Note that the entire set of nuclear parameters, not only the incompressibility and the symmetry energy, but also their density dependence determine the resulting EOS. In this paper we present results of numerical simulations of core-collapse supernovae of massive iron-core progenitors. The spectral neutrino transport is treated by the {\it Isotropic Diffusion Source Approximation} (IDSA) \citep{lieb09}. We employ four EOS of LS (with the three incompressibilities) and SHEN. We apply these different EOS in axially symmetric (2D) simulations and investigate the differences obtained. In the analysis of the data we focus on the post-bounce phase after the shock-stall, on the neutrino-driven shock revival and the shock propagation for more than 500 ms after core bounce. The paper opens with the description of the the numerical method and the employed EOS in Section 2. The results of 1D and 2D simulations are presented in Section 3 and 4. We summarize our results and discuss their implications in Section 5.
|
\label{sec:summary} We performed 1D and 2D radiation hydrodynamic simulations that include spectral neutrino-radiation transfer. The simulations are launched from a 15 $M_\odot$ progenitor star using four different EOS, corresponding to three different incompressibilities of the LS EOS and one SHEN EOS. LS is based on a non-relativistic approach and has a symmetry energy of 29.3~MeV. The incompressibility parameter can be set to 180 (LS180), 220 (LS220), or 375 MeV (LS375). SHEN is based on relativistic mean field theory and the Thomas-Fermi approximation. It has a symmetry energy of 36.9 MeV and an incompressibility of 281 MeV. This set of of EOS represents a wide range of nuclear matter properties. In 1D, none of the simulations produced an explosion for the considered simulation times up to 1 s post bounce. The observed similarities between LS180, LS220 and LS375, as well as differences to SHEN, concerning the conditions at bounce, the propagation of the shock, and the evolution of neutrino luminosities and mean energies, are in agreement with previous studies. The incompressibility and symmetry energy of the nuclear EOS depend on the density and determine the conditions from which the explosion may emerge. However, as recently discussed in \cite{stei12}, it is not possible to reduce the characteristics of the EOS to a single important parameter. Note that there are several constraints on the nuclear properties \citep[e.g.,][]{stei10, latt12}. These suggest that, among the employed EOS in this study, SHEN has too high symmetry energy, LS180 has too low incompressibility, and LS375 has too high incompressibility. Only LS220 is compatible with terrestrial experiments and the 1.97 $M_\odot$ Demorest~\emph{et~al.} pulsar mass constraint for cold neutron star matter. In 2D, the differences between models with LS and SHEN EOS are consistent with the differences between the 1D models. For example, we find significantly more favorable conditions for the launch of a neutrino-driven explosion in the models using the LS EOS: The shock expands continuously to large radii, which are not reached in the models using SHEN at comparable times. We analyzed these different post-bounce evolutions and found that models with LS show a highly turbulent velocity field producing prompt convection during the early post-bounce phase. Moreover, LS leads to to higher electron antineutrino luminosity and more efficient neutrino heating. More mass is accumulated in the gain region, which leads to a higher entropy per baryon at the early post-bounce phase. In addition, the standing accretion shock instability (SASI) pushes the standing accretion shock to increasingly larger radii after a post-bounce time on the order of 100~ms. In contrast to these models based on LS, the models based on SHEN show weak neutrino heating and the standing accretion shock is only oscillating around a mean radius for simulation times up to 600~ms post bounce. Note, however, that our simulations ignore inelastic scattering on electrons, which is important during the collapse phase, and the emission of heavy-lepton neutrinos, which contributes to the energy loss during the post-bounce phase. These reactions would both make our models more pessimistic, so that the 2D models based on SHEN appear unlikely to produce an explosion with the complete set of neutrino reactions. Hence, the different properties of the nuclear EOS can explain some apparently contradictory results found in previous studies, for example the outcome of neutrino-driven explosions in simulations of \cite{bura06,brue09,suwa10} using LS EOS and the absence of neutrino-driven explosions in simulations of \cite{burr06} using SHEN. Very recently, \cite{couc12} investigated the EOS dependence based on simulations with parameterized neutirno luminosities similar to \cite{murp08,nord10,hank12}. He found that a 15 $M_{\odot}$ star is more difficult to explode using SHEN than LS EOS, which is also consistent with the result of this paper. The most optimistic model, with respect to a possible neutrino-driven explosion, was obtained using LS180, for which we explore different indicators for the possible onset of an explosion. We note again that the neutrino-driven shock revival occurres much faster in this model compared to \citet{mare09} because of the aforementioned neglect of the additional cooling processes. The energy for the shock expansion stems from aspherical mass accretion from large radii deep onto the surface of the protoneutron star, generating a semi-stationary advection cycle. Thus, matter which expands behind the shock wave experiences a continuous flux of neutrinos from the accreted matter. It remains to be shown whether the asymmetry of the accretion pattern is similarly pronounced in three-dimensional supernova simulations. A small difference between the $\nu_e$ and $\bar\nu_e$ spectra leads to proton-rich conditions. Whether these findings will remain after having included a treatment of charged-current processes that is consistent with the EOS, as discussed recently in \cite{mart12} and \cite{robe12}, remains to be shown in future studies. Such improvements may become important especially at later post-bounce times that we have not reached in the present simulations. Based on the described explosion model we evaluate different indicators that have been suggested in the literature to diagnose the onset of an explosion during the post-bounce phase. We confirm that mass-weighted average entropy in the gain region \citep{murp08} and the mass enclosed in the gain region itself \citep{jank01} both show the expected difference between the optimistic LS and the pessimistic SHEN models. With respect to the ante-sonic condition \citep{pejc12}, we find that it applies as well. However, fast convection can lead to transient peaks in the indicator that may not be interpreted as a possible onset of the explosion. Also the ratio between advection time scale and heating time scale has been suggested as explosion indicator \citep{thomp05}. This indicator is also affected by multidimensional effects and one would have to carefully specify how the time scales have to be averaged in a multidimensional setting. The evaluation of the time scales based on angularly integrated quantities does not lead to a robust indicator for the onset of an explosion. Also the evolution of the radius of the protoneutron star does not serve as a reliable predictor of the explosion. In order to extend our EOS comparison study to a broader range of initial progenitor models, and in particular to confirm the optimistic results obtained using LS, we also include a lower mass iron-core progenitor of 11.2 $M_\odot$ into our investigation. For this model, we obtain neutrino-driven explosions for both LS and SHEN EOS. The differences observed between both simulations remain qualitatively similar to the ones discussed above for the 15 $M_\odot$ models. In comparison to LS, SHEN leads to a slightly delayed onset of the explosion with a smaller explosion energy. However, it should be noted that the simulations in this paper are only a very first step towards more realistic supernova models. In addition to the simplifications of weak processes and the omission of heavy-lepton neutrinos, general relativistic effects should be considered. Additionally, the ray-by-ray approximation may lead to an overestimation of the directional dependence of neutrino anisotropies and of radial oscillations of SASI. A full-angle transport will give us a more refined answer \citep[see][]{ott08,bran11,sumi12}. Moreover, due to the coordinate symmetry axis, the SASI develops preferentially along this axis. It could thus provide more favorable condition for the growth of $\ell = 1$ mode of the SASI and also for the possible onset of explosions along this direction. In the appendix we point to the clear and crucial impact the angular width of the computational domain has on the activity of the SASI, and in consequence on the presence or absence of the explosion. Hence, supernova modelers are gradually switching gears from 2D to 3D simulations \citep[e.g.,][]{iwak08,sche08,kota09,sche10,nord10,wang10,taki12,hank12,burrows12}, in which the SASI and convection were observed to develop much more stochastically in 3D than in 2D. This shows that complementary 3D supernova models are indeed necessary to pin down the impact of the EOS on the neutrino-driven mechanism. Together with the 3D effects, it would also be interesting to study the possible EOS impacts on the gravitational-wave signature and neutrino emission (e.g., \citealt{mare09b}, \citealt{muel12}, see \citealt{kota12} for a recent review). We will address these questions in forthcoming papers.
| 12
| 6
|
1206.6101
|
1206
|
1206.1384_arXiv.txt
|
The proof of cosmic ray (CR) origin in supernova remnants (SNR) must hinge on full consistency of the CR acceleration theory with the observations; direct proof is impossible because of the orbit stochasticity of CR particles. Recent observations of a number of galactic SNR strongly support the SNR-CR connection in general and the Fermi mechanism of CR acceleration, in particular. However, many SNR expand into weakly ionized dense gases, and so a significant revision of the mechanism is required to fit the data. We argue that strong ion-neutral collisions in the remnant surrounding lead to the steepening of the energy spectrum of accelerated particles by \emph{exactly one power}. The spectral break is caused by a partial evanescence of Alfven waves that confine particles to the accelerator. The gamma-ray spectrum generated in collisions of the accelerated protons with the ambient gas is also calculated. Using the recent Fermi spacecraft observation of the SNR W44 as an example, we demonstrate that the parent proton spectrum is a classical test particle power law $\propto E^{-2}$, steepening to $E^{-3}$ at $E_{br}\approx7GeV$.
|
The discovery of cosmic rays (CR) dates back to the historic Victor Hess balloon ascent in 1912 \cite{Hess12}. CR origin is thus a century old problem. Only the latest \emph{direct }observations of galactic supernova remnants (SNRs) \cite{Enomoto02,Ahar06RXJ,Abdo10W44full,VeritasCasA10,AbdoW28_10,AbdoIC443_10,Abdo11RXJ,AgileW44_11} narrowed the search to precisely these objects as the most probable sources of the CRs. One serious problem on the observational side was the lack of the SNR gamma-ray data below the energy range of the imaging atmospheric Cerenkov telescopes, or IACT. The Fermi-LAT (large area telescope) and Agile observatories are rapidly bridging this gap (roughly in the 0.1-30 GeV band e.g., \cite{Abdo10W44full,AgileW44_11}), virtually overlapping with the IACT energy band. There have been recent breakthrough observations of such SNR as W44, IC443, W28, RX J1713 and Cas A \cite{Abdo10W44full,AbdoIC443_10,AbdoCasA10,AbdoW28_10,Abdo11RXJ}. Overall, observations favor the diffusive shock acceleration (or DSA \cite{Drury83,BlandEich87,MDru01}, a modern version of the mechanism originally suggested by Fermi in 1949 \cite{Fermi49}) as a means for the production of galactic CRs. However, there are questions, and even some challenges, that the recent observations pose to the theory. Of those, the most relevant to the proof of the SNR-CR connection is the form of the spectrum that the theory predicts for the particular SNR conditions. Full understanding of the spectra will allow one to disentangle the proton (i.e., the primary CR component) emission from a contaminating (1-2\% level) but radiatively more efficient, and accessible to the direct observations, electron CR component. The most recent challenge to the DSA was posed by the measurements of the rigidity (momentum to charge) spectra of different species (most notably proton and helium). They turned out to be different, contrary to the DSA predictions for the ultra-relativistic rigidity range. Note that the latter problem arose from the indirect observations of the background CRs \cite{Adriani11,CREAM10,ATIC2_09,AMS_prot_00}, as opposed to the above mentioned direct observations of the putative accelerators (SNR). Generally, it is impossible to trace CR back to their accelerators because of the orbit scrambling. The proof of their origin in SNRs can only be achieved by proving the acceleration theory consistent with all accessible observations. It should be noted that 'direct' observations also provide only the secondary photon emission generated by accelerated particles, either electrons (through synchrotron, Bremsstrahlung and inverse Compton radiation), or protons (through their collisions with the ambient gas material). Therefore, such observations cannot be interpreted as an evidence of proton acceleration in SNR without a detailed understanding of the emission mechanism. Note that electron acceleration in SNRs to at least $\sim100$ TeV is held proven ``beyond a reasonable doubt'' after the observations of the SNR 1006 by ASCA and other X-ray instruments \cite{Koyama95,Allen01}. This paper deals with the modification of the DSA proton spectra in a partially ionized SNR environment and its signatures in gamma-emission from such remnants. The recent discovery of the proton/helium anomaly in the background CR spectra is discussed elsewhere (\cite{MDSPamela12}, see also \cite{VladimirMoskPamela11,Ohira11,BlasiChemComp11} for other suggestions to explain this anomaly). Here we pursue an alternative, complementary approach to more common multi-band treatments, e.g., \cite{TanakaRXJ08}, where the fits are primarily focused on the overall agreement across the entire spectrum (from radio to gamma). By contrast, we concentrate on the gamma-ray band and fit an important signature of the spectrum which is the spectral break. We believe it conveys an important information about the physics of acceleration missed in the 'standard' DSA theory. The quality of our fit, with virtually no adjustable parameters, should testify for the underlying physical scenario behind the emission. The broad-band fits do not typically meet high-quality criteria, as they seek to fit several portions of the data simultaneously by adjusting, in some cases, a few free parameters. Nevertheless, they provide an excellent consistency check for each particular model. The recent Fermi-LAT observations of the SNRs W44 and IC443 \cite{Abdo10W44full,AbdoIC443_10} indicate that the spectrum of the gamma ray producing protons is substantially steeper in its high energy part than the DSA predicts. A similar discrepancy has been found in the high energy gamma ray spectra measured by e.g., the CANGAROO \cite{Enomoto02}, HESS \cite{Ahar06RXJ} and MAGIC \cite{MAGICW51C11} atmospheric Cerenkov telescopes. The lack of understanding of the primary particle spectra triggered debates about the nature of the observed gamma-ray emission (hadronic vs leptonic), e.g., \cite{Pohl02}. We argued \cite{MDS05} that when a SNR interacts with a dense molecular cloud complex, the conditions for particle confinement to the shock are different from those adopted in conventional DSA modeling. Since the propagation of resonant Alfven waves is inhibited by ion-neutral collisions, particles are not confined and so escape the emission volume. These phenomena should result in a \emph{spectral break} in the parent proton and thus, in the gamma-ray spectrum. The spectral index at the break should change by exactly one power $\Delta q=1$ due to an effective reduction of particle momentum space dimension by one, since particles are confined in coordinate space only when they are within a slab in momentum space oriented perpendicular to the local mean magnetic field. Note that the earlier HESS observations of the SNR RXJ 1713 were also consistent with such a break \cite{Ahar06RXJ}. The most convincing evidence for the breaks of index one, however, provide the recent Fermi-LAT and Agile observations of W44 \cite{Abdo10W44full,Uchiyama10,AgileW44_11} (re-analyzed in \cite{MDS_11NatCo}), the MAGIC observations of the SNR W51C \cite{MAGICW51C11,MagicW51_12} as well as the FERMI observations \cite{Neronov12} of giant molecular clouds (GMC), where the Alfven wave evanescence should also result in a $\Delta q\simeq1$ steepening of the $E^{-q}$ CR primary spectrum. These observations are encouraging in that they unambiguously confirm the breaks. However, they rule out traditional DSA models based on a single power law with an exponential cutoff.
|
} To summarize the results, the mechanism for a break in the spectrum of shock accelerated protons suggested in \cite{MDS05,MDS_11NatCo} is in excellent agreement with the recent \cite{Abdo10W44full} Fermi LAT and Agile \cite{AgileW44_11} observations of the SNR W44. The observed gamma ray spectrum most likely results from the decay of $\pi^{0}$-mesons which are born in $p-p$ collisions of shock accelerated protons with an ambient dense gas. The parent proton spectrum is best represented by a classical test particle power law $\propto E^{-2}$, steepening to $E^{-3}$ at $E_{br}\approx7GeV$ due to deteriorated particle confinement caused by the ion-neutral collisions and the resultant Alfven wave evanescence. The position of the break momentum in the particle spectrum may be estimated using eq.(\ref{eq:p1}), or conversely, the combination of parameters involved in this estimate can be inferred from the measured break momentum. The cut-off momentum is not constrained in this scenario. An alternative explanation, based on a different mechanism of the break, associated with the change of the particle transport in the CR shock precursor \cite{MD06} is also possible but is less definitive in the spectrum slope variation $\Delta q$ across the break (see also \cite{Uchiyama10} for the most recent alternative suggestions). In addition, the mechanism \cite{MD06} would imply a considerable nonlinearity, i.e. a stronger CR shock precompression than that suggested by the radio observation of accelerated electrons and the inferred 100 GeV proton upper cutoff (see below). Still alternatively, assuming the ``environmental'' break mechanism is at work, i.e. $\Delta q=1$, but the shock structure is somewhat modified, we arrive at the $E^{-1.75}$ spectrum below the break (as the radio observations may suggest for the electrons), and $E^{-2.75}$ above the break. A fit to the data is marginally possible, but it would require a relatively low cut-off momentum at about $100$ GeV/c. This possibility may be supported or ruled out once the data (upper limit) around this energy become available. As we noted, particle escape from the MC can quench the acceleration process \cite{DruryEscape11}. This would certainly be the case if the MC were filling the entire shock precursor. However, MCs are known to be clumpy \cite{CrutcherMC99,Chev03,Pariz04,Inoue12}, and fill only a small fraction (< 1-2\%) of the precursor. In this case the acceleration process continues largely unimpeded (apart from the spectrum steepening) but the accelerated protons illuminate the 'cloudlets' and make them visible in $\gamma$-rays due to the high density target material. Another concern is a faint or even lacking x-ray emission that seems to be inconsistent with shocks impacting dense surroundings. This issue has been recently dealt with in, e.g., \cite{DruryHeating09,Inoue12}. Large clumps survive the shock passage as it stalls inside them and no strong heating occurs \cite{Inoue12}. The most robust and attractive aspect of the suggested mechanism for the spectral break is the exact $\Delta q=1$ variation of the spectral index. Indeed, this change in the spectral slope is due to the reduction of the number of degrees of freedom of particles caused by the resonant wave evanescence and it does not depend on any parameters. In a combination with the test particle regime operating below the break, which is physically suggested by the low values of the break and upper cut-off momenta, the mechanism provides a very good fit to the \emph{Fermi }LAT and Agile data with no free parameters for the SNR W44 and probably for W51C. From a number of physically different types of spectral breaks suggested \cite{AharAt96,mdj02,MD06,OhiraMC_11}, namely the current, ``environmental'' mechanism appears to be plausible where a dense target gas is present which is also required for the efficient $\pi^{0}$ production. However, observations of some other remnants in the dense gas environments, such as W28 and IC443 \cite{AbdoW28_10,AbdoIC443_10} indicate weaker breaks, $\Delta q=0.6-0.7$ which may either require a different mechanism for the break or a narrower wave evanescence gap $\Delta p=p_{2}-p_{1}$ (higher ionization rate). The predominance of small clumps with $L_{{\rm c}}\ll L_{{\rm CR}}$ in a MC will also reduce $\Delta q$. Generally, spectral breaks offer a natural resolution to the well known but puzzling trend of the \emph{nonlinear }(i.e. supposedly improved) DSA theory to develop spectra which are considerably harder than a simple test particle spectrum, thus becoming even less consistent with the bulk of observations \cite{Gaisser98,Hillas05}. However, the nonlinear spectrum -- i.e., diverging in energy-- exhausts the shock energy available for the acceleration as the cut-off momentum grows, so that a broken spectrum should form \cite{mdj02,MD06}. Broken spectra are now commonly observed and the old paradigm of a single power-law with an exponential upper cut-off is maladapted to the recent, revolutionarily improved observations \cite{AbdoIC443_10,Abdo10W44full}. Note, that the spectrum of the RX J1713.7-3946 \cite{Ahar06RXJ} is also consistent with the environmental break mechanism presumably operating in W44 surrounding but with a higher $p_{br}\sim10^{3}GeV/c$ and thus with stronger acceleration nonlinearity \cite{MDS05}. However, it is difficult to make the case for hadronic origin of the gamma-ray emission of the RX J1713.7-3946 \cite{AharNat04,Ahar06RXJ,WaxmanRXJ08}. The fundamental role of the W44 remnant for the problem of CR origin is that this particular remnant seems to rule out contaminating electron emission due to Bremsstrahlung and inverse Compton scattering \cite{Abdo10W44full,Uchiyama10} thus favoring the hadronic origin of the gamma emission and bolstering the case for the SNR origin of galactic CRs
| 12
| 6
|
1206.1384
|
1206
|
1206.1451_arXiv.txt
|
An essential role in the asteroidal dynamics is played by the mean motion resonances. Two-body planet-asteroid resonances are widely known, due to the Kirkwood gaps. Besides, so-called three-body mean motion resonances exist, in which an asteroid and two planets participate. Identification of asteroids in three-body (namely, Jupiter-Saturn-asteroid) resonances was initially accomplished by D.\,Nesvorn\'y and A.\,Morbidelli (1998), who, by means of visual analysis of the time behaviour of resonant arguments, found 255 asteroids to reside in such resonances. We develop specialized algorithms and software for massive automatic identification of asteroids in the three-body, as well as two-body, resonances of arbitrary order, by means of automatic analysis of the time behaviour of resonant arguments. In the computation of orbits, all essential perturbations are taken into account. We integrate the asteroidal orbits on the time interval of 100000~yr and identify main-belt asteroids in the three-body Jupiter-Saturn-asteroid resonances up to the 6th order inclusive, and in the two-body Jupiter-asteroid resonances up to the 9th order inclusive, in the set of $\sim 250000$ objects from the ``Asteroids -- Dynamic Site'' (AstDyS) database. The percentages of resonant objects, including extrapolations for higher-order resonances, are determined. In particular, the observed fraction of pure-resonant asteroids (those exhibiting resonant libration on the whole interval of integration) in the three-body resonances up to the 6th order inclusive is $\approx 0.9\%$ of the whole set; and, using a higher-order extrapolation, the actual total fraction of pure-resonant asteroids in the three-body resonances of all orders is estimated as $\approx 1.1\%$ of the whole set.
|
A substantial role of resonances in the dynamics of asteroids became evident with the discovery of resonant ``gaps'' in the asteroid belt by D.\,Kirkwood in 1867. The deepest minima in the distribution of asteroids in the semimajor axes of their orbits correspond to the mean motion resonances 2/1, 3/1, 4/1, 5/2, and 7/3 with Jupiter. Mean motion resonance represents a commensurability between the mean frequencies of the orbital motions of an asteroid and a planet. Apart from the mean motion resonances, so-called secular resonances \citep{MD99,M02}, representing commensurabilities between the precession rates of the orbits of an asteroid and a planet, are important in forming the dynamical structure of the asteroid belt. There are two important classes of the mean motion resonances: apart from the usual (two-body) mean motion resonances of an asteroid and a planet, an appreciable role in the asteroidal dynamics is played by so-called three-body mean motion resonances \citep{MHP98,NM98,NM99,M02}. In the latter case, the resonance represents a commensurability between the mean frequencies of the orbital motions of an asteroid and two planets (e.g., Jupiter and Saturn): \begin{equation} \label{dr} m_\mathrm{J}\dot{\lambda_\mathrm{J}}+m_\mathrm{S}\dot{\lambda_\mathrm{S}}+ m\dot{\lambda} \approx 0, \end{equation} \noindent where $\dot{\lambda_\mathrm{J}}$, $\dot{\lambda_\mathrm{S}}$, $\dot{\lambda}$ are the time derivatives of the mean longitudes of Jupiter, Saturn, and asteroid, respectively, and $m_\mathrm{J}$, $m_\mathrm{S}$, $m$ are integers. In view of the ``overdensity'' of the three-body resonances in the phase space of the asteroidal motion, \cite{NM98} asserted that ``the three-body mean motion resonances seem to be the main actors structuring the dynamics in the main asteroid belt''. Chaotic behaviour, which is often present in the dynamics of celestial bodies, is usually due to interaction of resonances (as in any Hamiltonian system, see \citealt{C79}), but not always it is known which are the interacting resonances that give rise to chaos. It is especially difficult to identify three-body resonances. How to distinguish between resonant and non-resonant motions? To solve this problem, a ``resonant argument'' (synonymously ``resonant phase'' or ``critical argument'') is introduced. It is a linear combination of some angular variables of a system under consideration; in the planar asteroidal problem it is given by \begin{equation} \label{radinition} \sigma_{p_\mathrm{J}, p_\mathrm{S}, p}=m_\mathrm{J}\lambda_\mathrm{J}+ m_\mathrm{S}\lambda_\mathrm{S}+m\lambda+p_\mathrm{J}\varpi_\mathrm{J}+ p_\mathrm{S}\varpi_\mathrm{S}+p\varpi, \end{equation} \noindent where $\lambda_\mathrm{J}$, $\lambda_\mathrm{S}$, $\lambda$, $\varpi_\mathrm{J}$, $\varpi_\mathrm{S}$, $\varpi$ are the mean longitudes and longitudes of perihelia of Jupiter, Saturn, and an asteroid, respectively, and $m_\mathrm{J}$, $m_\mathrm{S}$, $m$, $p_\mathrm{J}$, $p_\mathrm{S}$, $p$ are integers satisfying the D'Alembert rule \citep{M02}: \begin{equation} \label{dalambert_rule} m_\mathrm{J} + m_\mathrm{S} + m + p_\mathrm{J} + p_\mathrm{S} + p = 0 . \end{equation} If resonant argument~(\ref{radinition}) librates (similarly to librations of a pendulum), the system is in resonance; if it circulates, the system is out of resonance. The motion of the system at the border between librations and rotations corresponds to the {\it separatrix}. Thus the pendulum dynamics provides a graphical model of resonance. In a certain sense this model of resonance is ``universal'' \citep{C79}. In particular, the motion in three-body resonances can be described in the perturbed pendulum model \citep{MHP98, NM98, NM99, S07}. An important parameter of a mean motion resonance is its {\it order} $q$, equal to the absolute value of the algebraic sum of the coefficients at the mean longitudes in the resonant argument: \begin{equation} q = |m_\mathrm{J} + m_\mathrm{S} + m| . \end{equation} \noindent The resonant order $q$ is important, because it is the power in which the eccentricity is raised in the coefficient of the leading resonant term in the expansion of the perturbing function~\citep{NM98}. The corresponding subresonance width (characterizing also its ``strength'') is proportional to the square root of this coefficient. Thus the value of $q$ determines this important property of the leading subresonance. In the case of two-body resonances, the role of the resonant order $q$ (defined below in Section~\ref{sec_2br}) is analogous: the coefficient of the leading resonant term is proportional to $e^q$ \citep{NM98}, where $e$ is the asteroidal eccentricity. However note that, when there is no strong overlapping of subresonances, the resonant order $q$ is not related to the width of the whole resonant multiplet, because the separation of subresonances depends solely on the secular precession rates of the pericentres \citep{NM99}; thus the degree of overlap (and hence, chaos) in the multiplets is expected to asymptotically decrease with the resonant order \citep{NM99,MG97}. One may expect that, generally, broader the leading subresonance of a mean motion resonance, greater is the number of objects residing in this mean motion resonance. However, no strict correlation exists, due to a competition of various dynamical and physical processes, populating or depopulating the resonances. We shall discuss this further in more detail. In our procedure of resonance identification, described in detail below, we limit the set of possible combinations of the integers $m_\mathrm{J}$, $m_\mathrm{S}$, $m$ by adopting the following conditions: \begin{equation} \label{iac} q \leq q_\mathrm{max} , \end{equation} \begin{equation} \label{ic1} |m_\mathrm{J}|, \ |m_\mathrm{S}|, \ |m| \leq M_\mathrm{max}, \end{equation} \noindent where $q_\mathrm{max} = 6$ and $M_\mathrm{max} = 8$. \cite{NM98} used condition~(\ref{iac}) (presumably with $q_\mathrm{max} = 10$, as follows from data in table~3 in \citealt{NM98}). In \citep{NM99}, instead of (\ref{ic1}), the following truncation condition was used: \begin{equation} \label{inc} |m_\mathrm{J}|+|m_\mathrm{S}|+|m| \leq Q_\mathrm{max} \end{equation} \noindent (see eqs.~(29) and (30) and comments on them in \citealt{NM99}). We identify the three-body resonances in the current motion of asteroids with known orbital elements. The limitations of our study are as follows: solely the asteroids in the main belt are considered (i.e., the semimajor axes are in the range from 2 to 4~AU); solely the three-body resonances with Jupiter and Saturn are taken into account; the resonances are considered in the planar problem, i.e., the longitudes of nodes in the expression for the resonant argument are ignored; the maximum considered order $q_\mathrm{max}$ of the three-body resonances is set equal to 6. Our project is intended for the resonance analysis of the orbital data presented at the ``Asteroids -- Dynamic Site'' (AstDyS) maintained by A.\,Milani, Z.\,Kne\v{z}evi\'c and their coworkers \citep{AstDyS}. We take the orbital data for the analysis from this database. Thus the total set under analysis contains $\approx 250000$ objects. The basic purpose of our work is to identify the current three-body resonances that all the asteroids from the given set are currently involved in. More specifically, each object from the set should be put in correspondence to a three-body resonance (or none, if there is no resonance). The first attempt of massive identification of asteroids in three-body resonances was made by \cite{NM98}: 255 objects were identified to be in three-body resonances. The libration/circulation of the resonant argument for asteroids suspected to reside in the resonances was analyzed visually. In our case the data set is much greater, and therefore the procedure ought to be completely automatic. Besides, here we apply a unified bound on the order. This allows one to construct a homogeneous identification list for a further statistical analysis. To form a general statistical view of the resonant structure of the main belt, we also accomplish a massive identification of asteroids in two-body resonances with Jupiter, and compare the abundances of asteroids in three-body and two-body resonances.
|
\begin{enumerate} \item We have identified the resonant objects (the objects residing in three-body resonances with Jupiter and Saturn in the main asteroid belt) in the set of all numbered asteroids in the AstDyS database. This set comprises 249567 asteroids catalogued up to the date of April, 2011. The list of all asteroids identified as residing in pure three-body resonances is given in Appendix~A. \item The fraction of asteroids in three-body resonances (transient plus pure) up to the 6th order inclusive turns out to be $\approx 4.4\%$ of the total studied set of 249567 asteroids. The fraction of asteroids in pure three-body resonances of the same orders turns out to be $\approx 0.94\%$ of the total studied set. \item The top three most populated three-body resonances are: 5 -2 -2 (containing 699 transient+pure-resonant asteroids), 4 -2 -1 (688 transient+pure-resonant asteroids), 3 -2 -1 (621 transient+pure-resonant asteroids). For the pure-resonant asteroids, the ``top three'' resonances are: 4 -2 -1, 3 -1 -1, and 5 -2 -2, containing 595, 203, and 182 objects, respectively. \item Using a high-order extrapolation (in the form of a power law) of the $q$ dependence of the number of identified resonant objects, the actual total fraction of asteroids in pure three-body resonances of all orders is estimated as $\approx 1.1\%$ of the whole set. In what concerns the case of transient three-body resonances, the situation is much less certain, because the power-law extrapolation diverges. \item We have also identified all objects residing in two-body resonances (of order $0 \leq q \leq 9$) with Jupiter in the main asteroid belt, taking the same database of asteroids. The list of all asteroids identified as residing in pure two-body resonances is given in Appendix~B. \item The half of all identified asteroids in pure two-body resonances are Trojans ($\approx 53\%$). The pure Trojans plus pure Hildas constitute $\approx 85\%$ of all asteroids in the pure two-body resonances. The $q$ dependence of the two-body resonant abundances is clearly irregular and does not permit any smooth decay approximation. Especially one should point out the negligible asteroidal abundances at $q=3$ and $q=9$. \item In the transient plus pure resonances, the identified asteroids are $\approx 2.5$ times more abundant in the three-body resonances than in the two-body resonances; and in the pure resonances the abundances are comparable. However, if one excludes Trojans and Hildas, the abundance of three-body-resonant asteroids becomes overwhelming. What is more, taking into account extrapolated abundances in high-order resonances may substantially increase this overwhelming domination. Thus our analysis quantitatively verifies the assertion by \cite{NM98} that ``the three-body mean motion resonances seem to be the main actors structuring the dynamics in the main asteroid belt''. \item We would like to point out that our results confirm the general concept of \cite{M68,M69} on the omnipresence of resonances in the Solar system, however at a new level of understanding of this phenomenon. \end{enumerate}
| 12
| 6
|
1206.1451
|
1206
|
1206.5514_arXiv.txt
|
We report the discovery via radial velocity measurements of a short-period ($P = 2.430420 \pm 0.000006$ days) companion to the F-type main sequence star TYC 2930-00872-1. A long-term trend in the radial velocity data also suggests the presence of a tertiary stellar companion with $P > 2000$ days. High-resolution spectroscopy of the host star yields $T_{\rm eff} = 6427 \pm 33~{\rm K}$, $\log{g}=4.52 \pm 0.14$, and [Fe/H]=$-0.04\pm 0.05$. These parameters, combined with the broad-band spectral energy distribution and a parallax, allow us to infer a mass and radius of the host star of $M_1=1.21 \pm 0.08~\rm{M_\odot}$ and $R_1=1.09_{-0.13}^{+0.15}~\rm{R_\odot}$. The minimum mass of the inner companion is below the hydrogen burning limit, however the true mass is likely to be substantially higher. We are able to exclude transits of the inner companion with high confidence. Further, the host star spectrum exhibits a clear signature of Ca H and K core emission indicating stellar activity, but a lack of photometric variability and small $v\sin I$ suggest the primary's spin axis is oriented in a pole-on configuration. The rotational period of the primary estimated through an activity-rotation relation matches the orbital period of the inner companion to within $1.5\,\sigma$, suggesting that the primary and inner companion are tidally locked. If the inner companion's orbital angular momentum vector is aligned with the stellar spin axis as expected through tidal evolution, then it has a stellar mass of $\sim 0.3-0.4~\rm{M_\odot}$. Direct imaging limits the existence of stellar companions to projected separations $< 30$ AU. No set of spectral lines and no significant flux contribution to the spectral energy distribution from either companion are detected, which places individual upper mass limits of $M_{\left\{2,3\right\}} \lesssim 1.0 ~ \rm{M_{\odot}}$, provided they are not stellar remnants. If the tertiary is not a stellar remnant, then it likely has a mass of $\sim 0.5-0.6~\rm{M_\odot}$, and its orbit is likely significantly inclined from that of the secondary, suggesting that the Kozai-Lidov mechanism may have driven the dynamical evolution of this system.
|
Exoplanet surveys have contributed to a wide range of ancillary astrophysical disciplines during the last two decades, including studies of variable stars, binary stars and brown dwarf (BD) companions. During the course of operation, these surveys detect a large variety of stellar binaries that can be used to study stellar structure, atmospheres and formation mechanisms. One example of the latter is a study of the multiplicity of close binaries, e.g., the fraction of close binaries that are in triple or higher-order systems. Indeed, triple systems are not uncommon amongst short-period binaries; 9 out of 16 binaries with $P < 100$ days in the volume-limited sample of \citet{rag2010} are members of triple systems. Shorter-period binaries have an even greater probability of being in a multiple-star system \citep[$\sim 80$\% for $P < 7$ days vs. $\sim 40$\% for $P > 7$ days,][]{tok2006}. The orbital elements of such binaries, including the mutual inclinations of the companions' orbital angular momentum vectors, are fossil records of their formation process, and provide critical constraints to binary star formation models \citep{ste2002}. Comparison of the orbital and physical properties between different binary hierarchies also provides insight into binary star formation theory \citep{tok2008}. In fact, the dynamical evolution of these systems may be dominated by dynamical interactions between the inner and outer companions via a combination of the Kozai-Lidov mechanism \citep{koz1962,lid1962} and tidal forces, which drive the inner companion to shorter orbital separations until it circularizes with some period $P \lesssim 10$ days, beyond which tidal forces are ineffective \citep{fab2007}. In this paper, we present the discovery of a companion with a substellar minimum mass orbiting the bright ($V = 9.8$) F-type star TYC 2930-00872-1 \citep[][hereafter TYC 2930]{hog2000}, with an orbital period of $P = 2.430420 \pm 0.000006$ days. This discovery is part of a series of papers dedicated to analyses of individual low-mass companions in anticipation of a global analysis of the MARVELS sample at the conclusion of the survey \citep[e.g.,][]{lee2011,wis2012}, therefore, TYC 2930 is also designated ``MARVELS-2'' as an internal reference within this series. The a priori transit probability of the inner companion is $\sim$13\% with an expected central transit depth of ${\sim}0.9 \, \pm \, 0.25$\% for a $1 ~ \rm{R_{Jup}}$ companion radius, although no transits are detected. An additional, long-term trend in the RV data is detected from a stellar tertiary in the system. A detailed analysis of the combined radial velocity, spectroscopic and photometric data suggests the inner companion is oriented towards a pole-on configuration, and is more likely an M dwarf with a mass $\sim 0.3-0.4~\rm{M_{\odot}}$, while the tertiary is likely to be less inclined. In such a scenario, the mutual inclination between the secondary and tertiary is likely to be significant, which would make this an excellent example of a system whose dynamical history was driven via the Kozai-Lidov mechanism. The paper is organized such that \S \ref{specsec} describes the spectroscopic observations and their data processing, \S \ref{photobs} describes the archival and observed photometry for the system, \S \ref{starchar} describes the characterization of the host star's properties, including mass, radius, effective temperature, surface gravity, metallicity, stellar activity and rotation rate, \S \ref{orbitfit} describes our determination of the orbital parameters from fitting the measured RVs, \S \ref{imaging} describes both Lucky Imaging and adaptive optics imaging to search for any wide companions to TYC 2930, \S \ref{relphot} describes our search for photometric variability and any potential transits of the inner companion, \S \ref{geomtides} discusses the tidal evolution of the inner companion, \S \ref{massdistsection} describes the posterior distribution of the true masses for both the secondary and tertiary given the results from the previous sections, and finally, \S \ref{kozaisection} investigates the possible dynamical history of the system via the Kozai-Lidov mechanism.
|
We have discovered a short-period companion to TYC 2930-00872-1 with a minimum mass below the hydrogen burning limit. Despite its relatively high transit probability, we exclude any transits of the companion with high confidence using data from three ground-based telescopes. A long-term trend in the RVs indicate the presence of a longer-period tertiary in the system. The tertiary's spectral lines are not detected in our spectroscopic data, its fluxes do not significantly contribute to our SED fitting, and direct imaging excludes stellar-mass, main-sequence companions out to projected separations of 30 AU. Our spectra show the clear presence of Ca H and K core emission, but there is an unexpected lack of photometric variability, and the measured $v\sin I$ is significantly smaller than expected if the primary's rotation rate was tidally synchronized to the inner companion's orbital period. This suggests the primary's stellar spin axis is closely aligned to the line-of-sight. Given the age of the system, it is expected that the inner companion's orbital angular momentum vector is aligned with the stellar spin axis, therefore its line-of-sight orbital inclination is low, and its true mass is likely to be stellar. The absence of any detected signal from either component in the spectra and SED place an upper mass limit of $\sim 1.0 ~ \rm{M_{\odot}}$, if they are not stellar remnants. Assuming the tertiary is not a remnant, the upper mass limit places a lower limit to its line-of-sight inclination, which results in a significant mutual inclination between the secondary and tertiary. Such mutual inclinations are expected if the system's dynamical history was driven by the Kozai-Lidov mechanism. Long-term RV monitoring of the outer companion to obtain reliable orbital parameters will greatly improve the constraints that can be placed on the mutual inclination between the secondary and tertiary. Furthermore, high signal-to-noise ratio spectroscopic observations that could detect the presence of the (presumably) M dwarf companions would also allow for masses to be assigned to both objects and improve the inclination constraints.
| 12
| 6
|
1206.5514
|
1206
|
1206.2428_arXiv.txt
|
We study the evolution of an inflation-generated magnetic field, due to its coupling to fluid motions, during cosmological phase transitions. We find that the magnetic field stays almost unchanged on large scales, while on small scales the spectrum is modified in such a way that power at small scales becomes progressively suppressed. We also show that the magnetic field generates turbulent motions in the initially turbulence-free plasma. On large scales, the slope of the resulting kinetic energy spectrum is consistent with that of white noise.
|
The origin of the coherent large-scale ($\sim$ 10\,kpc) part of galactic magnetic fields, of $\mu$G strength, is under active discussion \cite{Subramanian:2010ab, Kandus:2011ab, Widrow:2012ab, Yamazaki:2012ab}. On larger, Mpc scales, until recently there were only upper limits, the most restrictive being of order a few nG depending on the observational technique used to measure the intergalactic magnetic field strength \cite{Kahniashvili:2010ab}. Recently, there have been a number of published lower limits on a putative large-scale magnetic field of strength $10^{-1 \pm 1}$\,fG (1\,fG = $10^{-15}$\,G) \citep{Neronov:2010ab}, or possibly two orders of magnitude smaller \citep{Dermer:2011ab}.\footnote{ These techniques for limiting a large scale cosmological magnetic field might be unreliable \cite{Broderick:2011ab}, but see their Sec.\ 4 where they note that more work will be needed to firm up these arguments and to determine whether the techniques used to establish the lower limits are indeed unreliable.} Almost certainly, the galactic fields are the amplified remnants of significantly weaker ``seed'' magnetic fields. Quantum mechanical fluctuations during inflation \cite{Fischler:1985ab} is a leading candidate for generating the needed seed magnetic field \cite{Turner:1988ab, Ratra:1992ab, Ratra:1991ab}. To generate a large enough seed magnetic field through quantum mechanical fluctuations during inflation, conformal invariance must be broken during inflation. A simple, realistic, illustrative model couples the abelian vector field with field strength tensor $F_{\mu\nu}$ to the scalar inflaton field $\phi$ through a dilaton-like coupling, generalizing the Maxwell lagrangian density to $e^{\alpha\phi} F_{\mu\nu} F^{\mu\nu}$ where $\alpha$ is a parameter \cite{Ratra:1992ab, Ratra:1991ab}. In the case of power-law inflation, and depending on the value of $\alpha$, this can result in a large enough seed magnetic field to explain the observed galactic magnetic fields. This is an observationally viable model. For a more detailed description of the model see Sec.\ II below. After the end of inflation, such an inflation-generated magnetic field will be correlated over super-Hubble-radius scales. It would induce observable signatures in the cosmic microwave background (CMB) radiation anisotropies at the epoch of recombination (the last scattering surface) if its current amplitude on Mpc scales is of the order of a nG \cite{CMBanisotropy}.\footnote{ The effects of an homogeneous magnetic field on the CMB anisotropy, and the resulting non-Gaussianity, are discussed in Refs.\ \cite{CMBnongaussianity}.} The properties of an inflation-generated primordial seed magnetic field depend on the parameters of the inflation model. If cosmological observations confirm the presence of an inflation-generated magnetic field, these measurements could be used to probe the physical conditions during inflation, including the shape of the inflaton potential energy density as well as the coupling between the inflaton and the vector gauge field. To check the consistency of the model, the primordial magnetic field shape and amplitude should be measured in as many ways as possible. The simplest limit arises from the cosmological expansion dynamics during big bang nucleosynthesis. This requires that the energy density of the magnetic field should not be larger than about 10\% of the radiation energy density. This limits the present (inflation-generated) magnetic field strength to less than a few $\mu$G, if the primordial magnetic field was generated prior to or during big bang nucleosynthesis, and was not damped or amplified by a magnetohydrodynamic (MHD) or some other process and so stays frozen into the plasma \cite{Kawasaki:2012ab}. In addition to the CMB temperature anisotropies that a primordial magnetic field induces (as mentioned above), such a field will Faraday-rotate the CMB polarization anisotropies \cite{Faraday}. Currently available Faraday rotation data give a bound on the primordial magnetic field strength of less than a few nG (for a scale-invariant or homogeneous primordial magnetic field). Another interesting signature of a cosmological magnetic field is the relic gravitational wave signal generated by the anisotropic magnetic stress \cite{Deriagin:1987ab}. The amplitude of the induced gravitational waves is determined by the magnetic field energy density, so a direct measurement of the resulting gravitational wave signal can lead to an independent limit on the magnetic field strength; see Ref.\ \cite{Wang:2010ab} and references therein. After the Universe reheats at the end of inflation, the plasma that was created then has large conductivity and it is conventional to assume that this remains the case as the Universe evolves to the present. In this case the large-scale cosmological magnetic field behaves as a frozen-in field with an evolution determined by the simple, flux-conservation, dilution of magnetic field lines, ${\bf B}({\bf x}, t) \propto {\bf B_0}({\bf x})/ a^2(t)$, where $t$ is the physical cosmic time and $a(t)$ is the cosmological scale factor. On the other hand, the evolution of a primordial magnetic field is a complex process influenced by MHD as well as by the dynamics of the Universe \cite{axel1,banerjee,campanelli2004, campanelli2007}. In particular, the presence of a magnetic field can dramatically affect primordial turbulence (e.g., when the turbulence is associated with cosmological phase transition bubble motions) \cite{axel1,axel-2,caprini09}. Furthermore, the presence of a magnetic field itself might lead to the development of turbulent motions, and so affect the turbulence \cite{B,axel}. In a recent examination \cite{kbtr10} of the effects of the MHD coupling between a primordial magnetic field and turbulence during a cosmological phase transition, we considered two different initial shapes for the spectrum of the primordial magnetic field: a single-scale magnetic field and a magnetic field with a Batchelor spectrum at large scales. In this paper we present a similar analysis for modified initial conditions for an inflation-generated primordial magnetic field \cite{Ratra:1992ab}, coupled via the usual MHD equations with the fluid, during the electroweak or QCD phase transitions. We consider both non-helical and helical magnetic field cases. We assume that the phase transition bubbles induce a typical length scale at which the magnetic field starts to interact with the phase transition fluid. The relevant difference between the electroweak and QCD phase transitions is encoded in the difference between values of parameters such as the temperature $T_\star$, the number of relativistic degrees of freedom $g_\star$, the bubble number, and bubble sizes. We assume initial absence of primordial turbulence, i.e., we assume that the plasma is initially at rest (although it is possible to generate turbulent motions through bubble collisions and nucleation \cite{kos}). The characteristic parameter of the primordial magnetic field is the r.m.s.\ Alfv\'en velocity $v_A = {B}/{\sqrt{16\pi {\rho_{\rm rad}}/3}}$. Here, $\rho_{\rm rad} \simeq \rho_{\rm thermal}$ is the radiation energy density. We use $\Omega_{\rm rad} h_0^2= 2.56 \times 10^{-5} $, where $\Omega_{\rm rad}$ is the radiation energy density parameter and $h_0$ is the Hubble constant in units of 100 km s$^{-1}$ Mpc$^{-1}$, for a current CMB temperature $T_0=2.74\,$K. At temperature $T_\star$, $\rho_{\rm rad}(T_\star) = {\pi^2}g_* (T_*)^4/30 $. The Alfv\'en velocity does not depend on $T_\star$ but is weakly dependent on $g_\star$, i.e., $v_A \propto g_\star^{-1/6}$. In our previous simulations \cite{kbtr10} we studied phase-transition-generated magnetic fields coupled to a relativistic fluid and discovered that equipartition between kinetic and magnetic energy densities is reached within reasonably short times. In this paper our main purpose is to consider different initial conditions. In particular, in the case when the magnetic field is generated during inflation, we investigate the kinds of turbulent motions that result from the coupling of the magnetic field with the fluid, and determine how this affects the evolution of the field itself. We show that the presence of a magnetic field on large scales ensures a rapid rise of the velocity field on large scales. On the other hand, magnetic field decay on large scales occurs at slow rates. The structure of our paper is as follows. In Sec.\ II we briefly describe magnetic field generation during inflation. In Sec.\ III we discuss the phenomenological coupling of the magnetic field to the turbulent plasma. In Sec.\ IV we present our numerical simulation results. We discuss our results and conclude in Sec.\ V.
|
In this paper we have studied the evolution of an inflation-generated magnetic field \cite{Ratra:1992ab} coupled to the fluid during cosmological phase transitions. Our formalism is very general and applies to the electroweak and QCD phase transitions. The difference between these (and other) phase transitions is encoded in the difference in parameters such as the temperature and the number of relativistic degrees of freedom, parameters which determine the characteristic length scale of the system under consideration ($\lambda_0$). We consider different types of forcing and show that at late times the kinetic energy spectrum depends sensitively on the forcing used. Our forcing scale is determined by the phase transition bubble size. Within a few turnover times the kinetic energy spectrum starts to rise on large scales, generating large-scale turbulent motions in the fluid. Even a rapid phase transition generates turbulence, which will slowly decay on large scales. Phase transition-generated MHD turbulence might be relevant for cosmological magnetogenesis \cite{Widrow:2012ab}. Phase transition turbulence can also generate a gravitational wave signal that is potentially detectable \cite{e-LISA}. In contrast to previous studies, the inflation-generated magnetic field is not frozen into the cosmic plasma. The forcing that we considered here is limited by the duration of the phase transition. After the forcing source stops to act, both magnetic and kinetic energies start to decay freely. The configuration of the magnetic field at large scales (outside the phase transition Hubble radius) is almost unchanged. At intermediate scales corresponding to the phase transition bubble size there is a slight suppression due to energy conversion into kinetic energy. The induced turbulent motions are causal so the spectral shape at large scales is given by a white noise spectrum $E_K(k) \propto k^2$ \cite{hogan}; the vorticity energy density spectrum will be steeper ($k^4$) due to the additional requirement of causality \cite{caprini2003}. The presence of magnetic helicity does not significantly change the forcing stage. On the other hand, the scaling laws in the decay stage are strongly affected by the presence of magnetic helicity. The duration of the decay stage is much longer than the forcing stage. During this stage the correlation length of the velocity increases with a corresponding decay of the total energy density. The magnetic field on super-Hubble radius scales is decoupled from the fluid which, in turn, stays almost unaffected. The main results of our study are: (i) inflation-generated magnetic fields are not significantly modified on large scales by their coupling to the plasma during a cosmological phase transition; (ii) the coupling of the magnetic field with the phase transition fluid leads to deviations of the magnetic field spectrum from the initial scale-invariant shape on intermediate scales; and, (iii) there is the possibility of having large-scale correlated turbulent motions in the early Universe which, eventually, could affect the development of large-scale structure formation at late times, and in particular cluster physics \cite{krav}.
| 12
| 6
|
1206.2428
|
1206
|
1206.0331_arXiv.txt
|
The advanced interferometer network will herald a new era in observational astronomy. There is a very strong science case to go beyond the advanced detector network and build detectors that operate in a frequency range from 1 Hz-10 kHz, with sensitivity a factor ten better in amplitude. Such detectors will be able to probe a range of topics in nuclear physics, astronomy, cosmology and fundamental physics, providing insights into many unsolved problems in these areas.
|
\ac{ET} is conceived to be a third generation detector whose conceptual design study was funded by the European Framework Programme FP7. The study completed in July 2011 helped produce a straw-man design of the detector and a summary of the science (both instrumental and astrophysical) that it promises to deliver \cite{DSD}. The accompanying article by Stefan Hild will discuss the technological challenges and the infrastructure needed for building ET. In this article we will discuss the rationale for going beyond advanced detectors and the huge spectrum of science and sources that ET has the potential to uncover. The discussion presented here is the result of a specific study carried out in the context of ET. However, much of it is relevant to an extension of either the advanced detectors or designs alternative to ET that target a sensitivity window from 1 Hz to 10 kHz, with the best strain sensitivity of $\sim {\rm few} \times 10^{-25}\,\rm Hz^{-1/2}$ in the frequency range of 20-200 Hz. What will ET observe in this frequency window? Why do we need detectors that are even more sensitive than the advanced detectors? What astrophysical problems can be addressed with ET? These are primary questions addressed in this article. ET, for that matter any \ac{GW} detector, is sensitive to compact objects with time-varying quadrupole moment. \acp{BH} and \acp{NS} being the most compact objects, close interactions between them, involving ultra-strong gravitational fields, will produce the most luminous gravitational radiation. ET's frequency range essentially determines the masses of compact objects that it could observe. The largest angular frequency which a BH of mass $M$ produces is roughly\footnote{We use a system of units in which the speed of light and gravitational constant are both equal to unity, $c=G=1.$ In this system, the mass, length and time have all the same dimensions, taken, for convenience, to be seconds.} $\omega^2 \sim M/R^3,$ where $R$ is its size. Taking $R=2M,$ the frequency works out to be $f\sim 1.14\,\, {\rm kHz}\, (M/10\, M_\odot)^{-1}.$ For comparison, the most dominant quasi-normal mode frequency of a $10\,M_\odot$ Schwarzschild BH is 1.19 kHz and that of a Kerr BH (with dimensionless spin of 0.9) is 2.15 kHz. Thus, the frequency range of 1-$10^4$ Hz gives a mass range of $1$-$10^4\,M_\odot.$ It might at first appear that the low-frequency window of 10-20 Hz, where the noise floor could be an order of magnitude or two larger than at 20-200 Hz, has no particular advantage for enhancing the visibility of signals. This is possibly true in the case of sources that sweep past the best part of the detector sensitivity. However, good low-frequency sensitivity does two things: Firstly, opens up a window for observing \acp{IMBH} \cite{Gair:2010dx,AmaroSeoane:2009ui} with masses in the range $10^3$-$10^4\,M_\odot.$ There is as yet no conclusive evidence for the existence of IMBH, let alone their binaries. However, there are strong indications that certain ultra-luminous x-ray sources (e.g.\, HLX-1 in ESO 243-49 \cite{2009Natur.460...73F}) are host to IMBH. If a population of such objects exists and they grow by merger, then, depending on their masses, ET will be able to explore their dynamics out to $z\sim6$-15 and study their mass function, redshift distribution and evolution. Secondly, lower frequencies help improve measurement accuracies of source parameters. Binary systems spend very long periods at lower frequencies, with the time to coalescence from a frequency $f$ rising as $f^{-8/3}.$ The long duration over which the sources slowly chirp-up in frequency helps in measuring the parameters of the source very accurately. For instance, in the case of advanced LIGO, the \ac{SNR} for a \ac{BNS} signal integrated from 10 Hz to 20 Hz is less than 1\% of the SNR that accumulates above 20 Hz until merger. Yet, the measurement accuracy of the system's masses is a factor of two better if the signal is integrated from 10 Hz instead of 20 Hz. This effect will be even stronger in the case of ET as its lower frequency cutoff could be a factor ten smaller compared to advanced detectors. The population of sources in the frequency window from 100 Hz to 10 kHz is also known to be very rich and there are many challenges and opportunities in this frequency region, both in instrument design and astrophysical potential. Quakes in \acp{NS} (believed to be the root cause of glitches in radio pulsar observations), giant explosions that occur in magnetars, gravitational collapse and supernovae, dynamics of accreting \acp{NS}, relativistic instabilities in young and accreting \acp{NS}, are all potential sources where observations could reveal a wealth of information that is complementary to radio, x-ray or gamma-ray observations. We will begin the discussion with a brief recap of what we can expect from a network of advanced detectors over the next decade. We will then go on to describe the topology and sensitivity of ET and why ET has additional advantages over equivalent L-shaped detectors. This is followed by a list of sources that ET can observe and how that benefits in furthering our knowledge of fundamental physics, cosmology and astrophysics. \begin{figure*}[t] \centering \includegraphics[width=24pc]{NetworkCumulDensity-DC.pdf} \caption{\label{fig:network rate}The plot shows the cumulative number of compact binary events expected to be detected by a network within a given distance, for three archetypal compact binaries and four different advanced detector networks. The curves flatten (and stay constant) upon reaching the {\em horizon distance} of the network, the distance beyond which a network cannot detect signals with the desired signal-to-noise ratios. See the text for further details. } \end{figure*}
| 12
| 6
|
1206.0331
|
|
1206
|
1206.4724_arXiv.txt
|
We study the dependence of the delayed neutrino-heating mechanism for core-collapse supernovae on the equation of state. Using a simplified treatment of the neutrino physics with a parameterized neutrino luminosity, we explore the relationship between explosion time, mass accretion rate, and neutrino luminosity for a 15 $M_\sun$ progenitor in 1D and 2D. We test three different equations of state commonly used in core-collapse simulations: the models of \citet{Lattimer:1991fz} with incompressibility of 180 MeV and 220 MeV, and the model of \citet{Shen:1998kx}, in order of increasing stiffness. We find that for a given neutrino luminosity the time after bounce until explosion increases with the stiffness of the equation of state: the Lattimer \& Swesty EOS explode more easily than that of Shen et al. We find this holds in both 1D and 2D, while for all models explosions are obtained more easily in 2D than in 1D. We also discuss the relevance of approximate instability criteria to realistic simulations.
|
\label{sec:intro} The exact process that halts the collapse of the core of a massive star at the end of its life and drives a successful supernova explosion, releasing a multitude of neutrinos and ejecta with around $10^{51}$ erg of kinetic energy is not yet fully-understood. The delayed neutrino-heating mechanism of core-collapse supernovae \citep[CCSNe,][]{Colgate:1966cl,Bethe:1985da}, the leading candidate for the explosion mechanism, fails to consistently produce energetic explosions for the wide range of possible progenitor masses in simulations that include high-fidelity treatments of neutrino transport. Neutrinos are an attractive explanation for the explosion mechanism as the bulk of the gravitational binding energy of the progenitor core, some $10^{53}$ erg, will be radiated away as neutrinos as the proto-neutron star (PNS) forms and cools; only about $10^{51}$ erg of this energy needs to be transferred to the collapsing stellar material in order to explain typical CCSNe energies. Only a fraction, however, of the core's original gravitational binding energy, perhaps a few times $10^{52}$ erg, will be released as neutrino radiation in the first second following the collapse and bounce of the core, approximately the timescale on which a successful explosion must occur. Neutrino radiation-hydrodynamic simulations of core collapse show that this is challenging to achieve, particularly in 1D simulations. It is now clear that multi-dimensional effects, such as proto-neutron star convection \citep{Epstein:1979tg,Burrows:1993ki,Dessart:2006cg,Burrows:2007kha}, neutrino-driven convection \citep{1996A&A...306..167J}, the standing accretion shock instability \citep[SASI,][]{Blondin:2003ep}, and turbulence \citep{Murphy:2011ci} play a critical role in the success of the neutrino mechanism in driving explosions. Together, these multidimensional effects can push some progenitors over the critical threshold, resulting in somewhat marginal explosions in 2D in certain simulations \citep{Marek:2007vi,Bruenn:2009cw,Suwa:2010wp,Mueller:2012tp}. This result is dependent on details of the numerical scheme and treatment of neutrino transport; the 2D simulations of \citet{Burrows:2006js,Burrows:2007kh} do not find neutrino-driven explosions for any progenitors.\footnote{ \citet{Burrows:2006js,Burrows:2007kh} find explosions via the ``acoustic'' mechanism first revealed by their simulations. The work of \citet{Weinberg:2008ky}, however, shows that the $\ell=1$ g-mode vibration of the neutron star which powers the acoustic mechanism will in reality saturate at energies well below that required to drive an explosion.} Based on a series of general relativistic 1D simulations including Boltzmann transport for neutrinos, \citet{Lentz:2012fy} suggest that the discrepancies in the results from the various groups simulating CCSNe may be due in large part to differences in how three major aspects of the physics are treated: general relativistic versus Newtonian dynamics, the neutrino interactions and opacities considered, particularly inclusion of inelastic neutrino scattering, and observer frame corrections in the neutrino transport. \begin{deluxetable*}{ccccccc} \tablecolumns{7} \tabletypesize{\scriptsize} \tablecaption{ EOS Parameters and experimental limits. \label{table:eosParams} } \tablewidth{0pt} \tablehead{ \colhead{EOS} & \colhead{$K$ \tablenotemark{a}} & \colhead{$K'$ \tablenotemark{b}} & \colhead{$J$ \tablenotemark{c}} & \colhead{$L$ \tablenotemark{d}} & \colhead{$M^{\rm max}_{\rm NS}$\ \tablenotemark{e}} & \colhead{$R_{1.4}$ \tablenotemark{f}} \\ \colhead{} & \colhead{(MeV)} & \colhead{(MeV)} & \colhead{(MeV)} & \colhead{(MeV)} & \colhead{($M_\sun$)} & \colhead{(km)} } \startdata LS180 & 180 & -451 & 28.6 & 74 & 1.84 & 12.2\\ LS220 & 220 & -411 & 28.6 & 74 & 2.06 & 12.7 \\ LS375 & 375 & -162 & 28.6 & 74 & 2.72 & 13.5 \\ STOS & 281 & -285 & 36.9 & 111 & 2.22 & 14.6 \\ HS (TMA) & 318 & -572 & 30.7 & 90 & 2.02 & 13.5 \\ & & & & \\ Limits & $240\pm10$ & -$355\pm95$ & $\sim 32$ & $\sim 75$ & $>1.97\pm0.04$ & 11 - 12 \enddata \tablenotetext{a}{Nuclear incompressibility.} \tablenotetext{b}{Skewness coefficient.} \tablenotetext{c}{Symmetry energy coefficient.} \tablenotetext{d}{Symmetry energy slope.} \tablenotetext{e}{Maximum gravitational mass of a cold neutron star.} \tablenotetext{f}{Radius of a 1.4 $M_\sun$ neutron star.} \end{deluxetable*} All of this indicates, unsurprisingly, that simulations of the neutrino mechanism are sensitive to microphysics, and details of how the microphysics is treated. It is still uncertain, however, if the variation in the handling of microphysics is large enough to explain the dearth of energetic explosions; that is, if a more accurate handling of the details of the neutrino physics will result in successful explosions having around $10^{51}$ erg of kinetic energy for a wide range of progenitor masses. \citet{Nordhaus:2010ct} have suggested a larger impact on the success of the neutrino mechanism results from fully 3D simulations. Using a simplified neutrino heating and cooling scheme rather than expensive neutrino transport, \citet{Nordhaus:2010ct} find that the driving neutrino luminosity necessary to obtain an explosion is about 20\% lower in 3D than in 2D. This difference may be larger than what we could expect to gain from higher-fidelity treatments of neutrino transport or inclusion of general relativity in simulations. \citet{Hanke:2011vc} sought to reproduce this result. Employing a neutrino heating/cooling approach similar to \citet{Nordhaus:2010ct}, \citet{Hanke:2011vc} find the critical neutrino luminosity in 3D is not significantly lower than that of 2D. This raises the questions of just how beneficial 3D simulations are to the success of the neutrino mechanism and what is the cause for the differences in the results of \citet{Nordhaus:2010ct} and \citet{Hanke:2011vc}? As 3D simulations become more feasible and available, we may be able to answer the question of the importance of 3D. Recently, \citet{Takiwaki:2012ck} report an explosion for an 11.2 $M_\sun$ progenitor in a 3D simulation performed using the isotropic diffusion source approximation for neutrinos \citep[IDSA,][]{Liebendorfer:2009kw}. Compared to 2D, Takiwaki et al. found that the convection below the gain region was more vigorous in 3D enhancing the neutrino luminosity beyond the 2D case. The somewhat modest resolution they were able to afford, however, means that the issue of the importance of 3D simulations cannot be settled on the basis of their results alone. \begin{figure} \centering \includegraphics[width=3.5in]{massRadius} \caption{Mass-Radius relations for cold neutron stars as computed by solving the Tolman-Oppenheimer-Volkov equations for the various EOS listed.} \label{fig:TOV} \end{figure} Besides the neutrino physics, an important microphysical dependence of the CCSN mechanism is the equation of state. The equation of state for matter at densities and temperatures relevant to CCSNe is still an active area of research \citep[e.g.,][]{Hempel:2010fh,Shen:2010wa,Shen:2011ta,Shen:2011wk,2011ApJS..197...20S,Furusawa:2011ck}. Our understanding of nuclear interactions at these densities and temperatures is incomplete and a number of models exist. Testing the various EOS models with experiment and observation is challenging, but a few constraints do exist. \citet{Hempel:2012bh} review some of the constraints on the EOS parameters from laboratory experiments. In Table \ref{table:eosParams} we summarize the constraints on the nuclear EOS parameters and the respective numbers for five EOS models used in CCSN simulations: Lattimer \& Swesty with incompressibility parameters of 180 MeV (LS180), 220 MeV (LS220), and 375 MeV (LS375), Shen et al. (STOS), and the recent EOS of \citet[][]{Hempel:2010fh} using the TMA model for nuclear interactions (HS (TMA)). Most values for the limits are taken from experiment \citep[for references to the experimental data see][]{Hempel:2012bh}, except for the skewness coefficient, $K'$, which is only a theoretical estimate. Based on the measurements listed in the table, only LS220 falls within or near the limits for each of the EOS parameters. Observations of neutron stars and pulsars can also place constraints on the EOS for hot dense matter. Using a parameterized EOS, \citet{Steiner:2010el} analyze a small sample of neutron stars and determine a most likely mass-radius relationship for cold neutron stars. Their results favor a radius of $11-12$ km for a 1.4 $M_\sun$ neutron star, and a maximum neutron star mass in the range $1.9-2.2$ $M_\sun$. The maximum mass of a cold neutron star is an important constraint for the EOS models. The observations of PSR J1614-2230 by \citet{Demorest:2010bf} place a very tight limit on the mass of the neutron star in this system of $1.97\pm0.04\ M_\sun$. Given an EOS model, solutions to the Tolman-Oppenheimer-Volkov equation yield the gravitational mass-radius relationship for cold neutron stars. Figure \ref{fig:TOV} shows the mass-radius relationships for the EOS of \citet{Lattimer:1991fz} and \citet{Shen:1998kx}, along with the mass limits for PSR J1614-2230. Based on the measurement of \citet{Demorest:2010bf}, LS180 is ruled out as it does not produce a large enough maximally-massed neutron star. The Shen et al. EOS produces a maximum neutron star mass well above the limit placed by PSR J1614-2230 (2.22 $M_\sun$), but STOS predicts a radius for a 1.4 $M_\sun$ NS of around 15 km, well outside the limits estimated by \citet{Steiner:2010el}. The theoretical predictions for the radius of a 1.4 $M_\sun$ NS of \citet{Hebeler:2010dx} are somewhat broader: 9.7 - 13.9 km, but STOS still lies outside of this range. Only LS220 satisfies the constraint on the maximum NS mass from PSR J1614-2230 and produces a 1.4 $M_\sun$ NS with a radius ($\sim 12.7$ km) close to the upper estimate of \citet{Steiner:2010el} and well within the limits of \citet{Hebeler:2010dx}. Figure \ref{fig:TOV} also shows the mass-radius relationship for LS375. This EOS easily satisfies the requirement on the maximum neutron star mass, but falls above the radius limits of \citet{Steiner:2010el} and has an incompressibility (375 MeV) well above that found in experiment. In this work, we study the dependence of the neutrino mechanism on the equation of state employed. Using 1D and 2D hydrodynamic simulations with simplified neutrino transport, we determine the explosion times for a 15 $M_\sun$ progenitor as a function of neutrino luminosity for three EOS: \citet{Lattimer:1991fz} with $K=180$ MeV and $K=220$ MeV, and \citet*[][]{Shen:1998kx}. In 1D only, we also run simulations using LS375. Our treatment of the neutrino physics is similar to those used by \citet{Murphy:2008ij}, \citet{Nordhaus:2010ct}, and \citet{Hanke:2011vc}: we assume local neutrino heating and cooling, based on the rates derived by \citet{Janka:2001fp}, and a constant neutrino luminosity. We find that the explosion times for a given neutrino luminosity are significantly dependent on the EOS used, with the general trend of easier explosions for softer equations of state. Thus, the Lattimer \& Swesty models result in easier explosions than the Shen et al. EOS, with LS180 leading to the earliest explosions at a given neutrino luminosity. Early 1D simulations using a simple, parameterized EOS showed that stronger shocks result for softer EOS \citep{Baron:1985ec, Baron:1985cd}. Our results show that this trend is recovered for more realistic nuclear EOS and in 2D simulations. The EOS has been shown to influence the results of CCSN simulations in a number of contexts. \citet{Thompson:2003kn} explored the sensitive of their 1D neutrino radiation hydrodynamic simulations to the value of $K$ for the Lattimer \& Swesty EOS. They found that the overall evolution during the first 200 ms post-bounce is not very dependent on $K$: the resulting temperatures were slightly higher and the emergent neutrino luminosities at most 9\% larger for LS180 as compared to LS375. \citet{Sumiyoshi:2005kg} compared LS180 and STOS in 1D general relativistic simulations. While neither model explodes in their simulations, they find differences in composition resulting from the different symmetry energies (see $J$ in Table \ref{table:eosParams}) and in temperature evolution with LS180 giving slightly larger temperatures and, therefore, slightly higher neutrino luminosities. The higher neutrino luminosity and heating rate found for LS180 is somewhat counteracted by the more compact proto-neutron star that results from the lower incompressibility. \citet{Sumiyoshi:2005kg} find that the differences between the two models grow with time, particularly after 300 ms post-bounce. \citet{Marek:2009bx} explored the dependence on the gravitational wave signal produced by 2D simulations on the EOS, comparing LS180 and the stiffer EOS of \citet{Hillebrandt:1985to}. They find that the matter-generated gravitational wave signal is not very dependent on the EOS but that the differences in the anisotropic neutrino emission gravitational wave signal, which can dominate the matter gravitational wave signal, are significantly different between LS180 and Hillebrandt \& Wolff EOS. In failed explosions, the formation of black holes has been found to be highly dependent on the EOS \citep{Sumiyoshi:2007cd, Fischer:2009ka, OConnor:2011hk, Hempel:2012bh}. Generally, stiffer EOS result in a greater time to black hole formation following bounce. This paper is organized as follows. In Section \ref{sec:Method} we describe the details of our numerical approach. In Section \ref{sec:Results} we describe the results of our simulations and the dependence of the explosions on the EOS. In Section \ref{sec:Conclusions} we conclude and discuss how our work relates to recent work on neutrino-driven supernovae.
|
\label{sec:Conclusions} We have conducted a parameter study exploring the dependence of the neutrino mechanism of core-collapse supernovae on the equation of state. We varied the driving neutrino luminosity and EOS in several 1D and 2D simulations of stellar core collapse and measured the resulting explosion times and mass accretion rates at the time of explosion. Table \ref{table:results} and Figure \ref{fig:massAcc} summarize our results. We find that for every EOS, 2D explosion are obtained more easily than in 1D, and that models using the Lattimer \& Swesty EOS explodes more easily than models using the EOS of Shen et al., with LS180 resulting in slightly earlier explosions than LS220. Thus, our results show a general trend that softer EOS lead to easier explosions. In order to test if the incompressibility parameter, $K$, is the primary determinant of the time delay to explosion for a given neutrino luminosity, we also ran a series of 1D simulations using LS375. This model has a substantially higher incompressibility than STOS (375 MeV v. 281 MeV). We find that STOS, despite have a lower incompressibility, explodes later than LS375 for all neutrino luminosities. This highlights that the higher-order EOS parameters, such as symmetry energy coefficient, play an important role in determining the stiffness of an EOS. Our results of explosion delay times do follow a trend in resulting neutron star radii (see Fig. \ref{fig:TOV}). EOS that produce larger typical neutron star radii explode later in our simulations. The dependence of the time-to-explosion on the EOS is likely due to the dependence of the acoustic energy flux from the PNS on the stiffness of the EOS. The softer EOS of Lattimer \& Swesty result in more compact PNS that are more effective at translating advected perturbations into outgoing sound waves. This then results in more acoustic excitation of the standing accretion shock instability and an overall faster rate of shock expansion. This should be true no matter if the underlying SASI mechanism is acoustic \citep{Blondin:2006dv} or advective-acoustic \citep{Foglizzo:2007cq} since both mechanisms depend on the acoustic flux from the PNS. For high-$L_{\nu_e}$ models that do not depend strongly on the SASI, we attribute the difference between STOS and LS to the higher mass accretion rate just after bounce for STOS. In this study, we use a parameterized neutrino ``light-bulb" with a fixed neutrino luminosity. There is no back-reaction from the EOS on the neutrino luminosity, allowing us to examine the dependence of the resulting explosion times on the EOS alone. Previous studies have shown that the neutrino luminosity can be significantly dependent on the EOS \citep{Thompson:2003kn, Sumiyoshi:2005kg} and this is a very important consequence of the choice of EOS, but our study shows that even for a fixed neutrino luminosity, the likelihood of a given model to explode also depends on the EOS. The EOS for matter at temperatures and densities relevant to CCSNe is still an active area of research. As summarized in the Introduction and Table \ref{table:eosParams}, available experiments and observations that constrain the EOS parameters favor LS220. LS180 is ruled out as it cannot produce a $\sim 2 M_\sun$ neutron star as required by the observations of PSR J1614-2230 \citep{Demorest:2010bf}. STOS yields neutron stars of typical mass (around 1.4$M_\sun$) with too large a radius ($\sim15$ km) to account for the results of \citet{Steiner:2010el} who estimate typical NS radii of 11-12 km. Several of the EOS parameters listed in Table \ref{table:eosParams} for STOS also fall outside of the experimental and theoretical estimates. So, LS220 is the EOS that satisfies, or nearly satisfies, the largest number of constraints. Our results show, additionally, that LS220 is intermediate between STOS and LS180 in favoring explosion for a given neutrino luminosity. Until we are more certain about the EOS at temperatures and densities relevant to CCSN simulations, LS220 seems a good choice. Some recent works have chosen to use LS220 on these same grounds \citep{Ott:2012ta, Dessart:2012vi}.
| 12
| 6
|
1206.4724
|
1206
|
1206.3797_arXiv.txt
|
We present {\it Hubble Space Telescope} imaging and spectroscopy, along with supporting GALEX and ground-based data, for the extended high-ionization cloud known as Hanny's Voorwerp, near the spiral galaxy IC 2497. WFC3 images show complex dust absorption near the nucleus of IC 2497. The galaxy core in these data is, within the errors, coincident with the VLBI core component marking the active nucleus. STIS optical spectra show the AGN to be a type 2 Seyfert of rather low luminosity. The derived ionization parameter log $U = -3.5$ is in accord with the weak X-ray emission from the AGN. We find no high-ionization gas near the nucleus, adding to the evidence that the AGN is currently at a low radiative output (perhaps with the central black hole having switched to a mode dominated by kinetic energy). The nucleus is accompanied by an expanding ring of ionized gas $\approx 500$ pc in projected diameter on the side opposite Hanny's Voorwerp. Where sampled by the STIS slit, this ring has Doppler offset $\approx 300$ km s$^{-1}$ from the nucleus, implying a kinematic age $< 7 \times 10^5$ years. Narrowband [O III] and H$\alpha$+[N II] ACS images show fine structure in Hanny's Voorwerp, including limb-brightened sections suggesting modest interaction with a galactic outflow and small areas where H$\alpha$ is strong. We identify these latter regions as regions ionized by recent star formation, in contrast to the AGN ionization of the entire cloud. These candidate ``normal" H II regions contain blue continuum objects, whose colors are consistent with young stellar populations; they appear only in a 2-kpc region toward IC 2497 in projection, perhaps meaning that the star formation was triggered by compression from a narrow outflow. The ionization-sensitive ratio [O III]/H$\alpha$ shows broad bands across the object at a skew angle to the galaxy nucleus, and no discernible pattern near the prominent ``hole" in the ionized gas. The independence of ionization and surface brightness suggests that there is substantial spatial structure which remains unresolved, to such an extent that the surface brightness samples the number of denser filaments rather than the characteristic density in emission regions; this might be a typical feature of gas in tidal tails, currently measurable only when such gas is highly ionized. These results fit with our picture of an ionization echo from an AGN whose ionizing luminosity has dropped by a factor $>100$ (and possibly much more) within the last $1-2 \times 10^5$ years; we suggest a tentative sequence of events in IC 2497 and discuss implications of such rapid fluctuations in luminosity for our understanding of AGN demographics.
|
The central energy sources of active galactic nuclei (AGN) are known to vary on a wide range of timescales. Direct observation samples variations from hours to decades (sometimes strong, and including dramatic changes in the prominence of the broad-line region). The dramatic evolution of luminous AGN with redshift demonstrates cosmological evolution, involving the entire AGN population. What this evolution entails for individual AGN is only indirectly constrained. Several arguments suggest that the central black holes grow episodically; the duty cycles, amount of accretion, or timescales for these episodes decrease with cosmic time (\citealt{Martini2004}, \citealt{Hopkins2005}). Luminous QSOs cannot maintain the observed energy output for much of cosmic history without the black holes becoming more massive than any we observe, and evidence for an excess of interaction signatures in QSO host galaxies (although not clearly present at lower luminosities) suggests that accretion is enhanced over roughly the timespan that we can recognize these signatures (typically a few times $10^8$ years). However, none of these factors reveals how the energy output behaves over timescales of $10^3$--$10^7$ years, spanning values for activity scales in quasars suggested by some observational considerations (\citealt{Martini2003}, \citealt{Kirkman}) as well as calculations of instabilities in accretion disks (\citealt{Shields1978}, \citealt{Goodman}, \citealt{Janiuk}, \citealt{Done}). Timescales of accretion-disk behavior may be estimated from scaling the state changes seen in X-ray binaries containing stellar-mass black holes (\citealt{Maccarone}, \citealt{McHardy}. We describe here new observations of an object which may hold key insights to otherwise unprobed timescales in the history of individual AGN - Hanny's Voorwerp. Among the signature serendipitous discoveries of the Galaxy Zoo project \citep{Lintott08}, this is a giant high-ionization nebula near the bright spiral galaxy IC 2497 \citep{Lintott09}. It was reported on the project forum\footnote{www.galaxyzooforum.org} by citizen scientist and co-author Hanny van Arkel, only a few weeks into the Galaxy Zoo examination of the SDSS main galaxy sample. Followup observations revealed a unique combination of characteristics, indicating that this object traces a luminous AGN which must be unusual either in obscuration or history. We first briefly summarize results from \cite{Lintott09}, \cite{Josza}, \cite{Rampadarath}, and \cite{Schawinski2010}. Hanny's Voorwerp is a region of highly-ionized material 18 by 33 kpc in projected extent, extending at least 50 kpc from IC 2497 and closely matching its redshift $z=0.050$. The electron temperature measured from [O III] lines indicates that it is photoionized rather than shock-ionized. Such emission-line ratios as He II/H$\beta$ and [Ne V]/[Ne III] show that the ionizing continuum is hard like an AGN rather than hot stars, while the ionizing luminosity for a source in IC 2497 must be of order $2 \times 10^{45}$ erg s$^{-1}$ to give the observed ionization parameter and intensity of recombination lines. However, the nucleus of IC 2497 has very modest luminosity, with emission lines implying ionizing luminosity $< 10^{40}$ erg s$^{-1}$. H I observations show that Hanny's Voorwerp is a small part of a 300-kpc structure around the southern side of IC 2497 containing $9 \times 10^9$ M$_{\odot}$ of neutral hydrogen. The nucleus hosts a compact VLBI radio source of modest power, and an additional feature which could be the brightest knot in a jet pointing roughly toward Hanny's Voorwerp. Lower-resolution radio continuum data show what may be a broad outflow in the same direction. The ionizing continuum required is of a luminosity associated with QSOs, making IC 2497 a very nearby QSO host galaxy observable in great detail. The spectra in \cite{Lintott09} indicate substantially subsolar abundances in Hanny's Voorwerp (and by extension in the H I tail). For gas photoionized by an AGN continuum across the relevant range of ionization parameter $U$, the [N II]/H$\alpha$ ratio is almost purely an abundance indicator. Models by \cite{Groves06}, \cite{FuStockton07}, and \cite{FuStockton} using the MAPPINGS code \citep{Groves04} show that this line ratio remains consistent for various ionization levels relevant to AGN narrow-line regions and extended regions. From Fig. 11 of \cite{FuStockton}, Hanny's Voorwerp has N/H $0.3 \pm 0.1$ solar. These data led to two competing interpretations - that the AGN is either hidden or faded. \cite{Lintott09} introduced the ionization-echo hypothesis, driven by the lack of detected X-rays from IC 2497, lack of any high-ionization gas observed in the galaxy, and the shortfall between the expected far-infrared luminosity and what is observed if most of the AGN output is absorbed by dust. \cite{Josza} favored a picture in which an outflow, seen in the radio continuum, cleared a path for ionizing radiation to escape circumnuclear obscuration, so that the Voorwerp would be ionized by an extant AGN which is so deeply obscured from our direction that not even the soft X-rays detectable by {\it Swift} would emerge. Evidence that QSOs can fade so rapidly would have a significant impact on our understanding of the demographics of QSOs. IC 2497 is much nearer than any known QSO of its inferred luminosity, yet has managed to elude all standard ways of surveying for them, and it is unlikely that a very rare event would be represented so close to us. This is likely to represent a phenomenon which is so common among AGN as to alter our estimates of their overall behavior. The key role of emerging X-rays in distinguishing between fading and obscuration motivated a set of {\it XMM-Newton} and {\it Suzaku} observations \citep{Schawinski2010}. The results support a fading scenario. The nuclear X-ray source in IC 2497 is well fitted by a combination of hot ISM and an AGN which is essentially unobscured beyond the Galactic H I value of $1.3 \times 10^{20}$ cm$^{-2}$. There is no detected 6.4-keV K$\alpha$ feature which characterizes deeply obscured ``reflection" AGN, and the {\it Suzaku} data above 10 keV in particular rule out a Compton-thick AGN luminous enough to ionize Hanny's Voorwerp. The 2-10 keV luminosity of $4.2 \times 10^{40}$ erg s$^{-1}$ falls short of the luminosity needed to account for the ionization of Hanny's Voorwerp, by four orders of magnitude. As part of our effort to unravel the nature of this system, we have obtained {\it Hubble Space Telescope} (HST) observations and supporting data, from the mid-ultraviolet to the near-infrared. These allow us to address the ionization structure of Hanny's Voorwerp, the ionization and kinematics of gas near the nucleus of IC 2497, and evidence for compression-induced star formation suggesting that an outflow from IC 2497 is interacting with a portion of Hanny's Voorwerp.
|
Our data support the scheme of a quasar rapidly dropping in luminosity; STIS spectra show no highly-ionized gas in IC 2497. However, the STIS spectra and images show signs of outflow from the nucleus, leading to the possibility that some of the energy output from the AGN has switched from radiation to kinetic form. This is seen in an expanding loop of gas $\approx 500 $ pc in diameter, and in star-forming regions within Hanny's Voorwerp which we attribute to compression by the outflow seen at radio wavelengths. We suggest the following sequence of events: A major merger liberated the massive tail of H I, with geometry such that the remnant IC 2497 retained its disk although significantly warping it. The low metallicity of the gas suggests that it began in the extreme outer disk of the galaxy. The dynamical disturbance eventually triggered an episode of accretion into the central black hole, with an ionizing luminosity appropriate for a quasar. The escaping UV radiation ionized parts of the H I tail, creating Hanny's Voorwerp. Recently (perhaps a few million years before our current view), an outflow from the core began, including the small-scale radio jet, the emission-line ring near the nucleus, and a narrow outflow directed roughly toward Hanny's Voorwerp, triggering star formation in the small region where this outflow impinged on relatively dense gas. Then within the last 100,000 years before our view of the galaxy nucleus, the ionizing luminosity dropped enormously, by 2--4 orders of magnitude, leaving Hanny's Voorwerp as the only remaining evidence of this episode. The close association of this drop in time with the onset of outflows may indicate that this was not completely a corresponding drop in the level of activity of the nucleus, but rather a switch between so-called quasar and radio modes of accretion. The light-travel delay across the system is small compared to limits on the age of the outflows, so that we cannot say whether this fading was a one-time event or recurs after multiple bright episodes. The massive H I tail is crucial in making the history of the AGN observable. Since such tails could be regarded as ``living fossils" of epochs when major mergers were common, we might expect to see more objects of this kind as data of the necessary spatial resolution become available at larger redshifts. If the combination of fading ionization and onset of outflows proves to be common, we might also find isolated star-forming regions near formerly active galaxies along the axis of an outflow, outlasting all but the very low-frequency radio emission associated with the outflow itself. This nearest quasar escaped all the standard optical, X-ray, and radio survey techniques. The fact that IC 2497 is at such a low redshift (and would have been the nearest AGN of its luminosity) would be very unlikely unless similar episodic behavior is common among AGN. Indeed, a dedicated search by Galaxy Zoo participants has found additional, less luminous examples of large-scale ionized clouds, including additional potential examples of fading AGN, in the local Universe \citep{Keel11}. These are found systematically in disturbed systems, attesting to the importance of an extended reservoir of neutral gas as a probe of the pattern and history of ionizing radiation from AGN. Additional surveys are in progress, keying on [O III] emission as a sensitive tracer of highly-ionized gas around samples of galaxies with and without detected AGN.
| 12
| 6
|
1206.3797
|
1206
|
1206.3568_arXiv.txt
|
We study the linear and nonlinear structure formation in the dilaton and symmetron models of modified gravity using a generic parameterisation which describes a large class of scenarios using only a few parameters, such as the coupling \g{between} the scalar field \g{and the matter,} and the range of the scalar force on very large scales. For this we have modified the $N$-body simulation code {\tt ECOSMOG}\g{, which is a variant of {\tt RAMSES} working in modified gravity scenarios,} to perform a set of $110$ simulations for different models and parameter values, including the default $\Lambda$CDM. These simulations enable us to explore a large portion of the parameter space. We have studied the effects of modified gravity on the matter power spectrum and mass function, and found a rich and interesting phenomenology where the difference with the $\Lambda$CDM template cannot be reproduced by a linear analysis even on scales as large as $k\sim0.05$ $h\rm{Mpc}^{-1}$. Our results show the full effect of screening on nonlinear structure formation and the associated deviation \g{from} $\Lambda$CDM. \BLED{We also investigate how differences in the \g{force mediated by the scalar field in} modified gravity models lead to qualitatively different features for the nonlinear power spectrum and the halo mass function, and how varying the individual model parameters changes these observables. The differences are particularly large in the nonlinear power spectra whose shapes for $f(R)$, dilaton and symmetron models vary greatly, and where the characteristic bump around $1\ h\rm{Mpc}^{-1}$ of $f(R)$ models is preserved for symmetrons, whereas an increase on much smaller scales is particular to symmetrons. No bump is present for dilatons where a flattening of the power spectrum takes place on small scales. These deviations from \g{$\Lambda$CDM} and the differences between modified gravity models, such as dilatons and symmetrons, could be tested with future surveys. }
|
\label{sect:introduction} The apparent acceleration of the Universe could be due to at least four different reasons: a cosmological constant, dark energy \citep{cst2006}, modified gravity \citep{cfps2011} or large spatial inhomogeneities \cite{biswas}. The last of these violates the Copernican principle and requires a theory for the initial conditions of the Universe while the first three invoke a change of the dynamics of the Universe itself. The cosmological constant solution is rather peculiar as no real dynamics is attached to it until the vacuum energy starts dominating the energy content of the Universe. This seems to have happened in the quite recent past, a fact which is problematic and related to the astoundingly small value of the critical density of the Universe compared to particle physics expectations, which scale as the fourth power of the mass of any heavy particle present in the early Universe. To alleviate this problem, two other possibilities are commonly invoked. The first one is dark energy \citep{cst2006}, in which the dynamics of a field \g{(\eg, a scalar field in the simplest case)} determines the \g{fate of the Universe}. So far no real solution to the cosmological constant problem has been found within this setting although phenomenological works abound. Setting aside the problem of the actual value of the dark energy density now, these models suffer from another serious problem: \HANS{dark energy evolves on cosmological time scales only when the scalar field leads to a long range interaction.} Of course, one can decree that dark energy does not couple to baryons as in coupled quintessence models\footnote{\BLED{We regard the coupled quintessence model as an example of dark energy rather than modified gravity, for which we require a universal coupling to all matter species.}}, and therefore alleviate gravitational problems linked to the existence of a scalar fifth force. If this is not the case, then a solution which has been put forward in the last decade is screened modified gravity mediated by a scalar field. Many models of \BLED{screened} modified gravity have been constructed so far, which fall within two broad categories. Following the initial works on massive gravity, models involving nonlinear kinetic terms, such as the Galileon \citep{dgp2000,nrt2009,dev2009}, make use of the Vainshtein mechanism \cite{vainshtein} whereby large nonlinearities in the vicinity of dense objects effectively reduce the scalar coupling to matter to be below the experimental bounds. Another class of models originating from the chameleon theory \citep{kw2004,ms2007} use a screening of the fifth force in dense environments due to the nonlinearities of either the scalar potential or its coupling to matter (or both). Chameleon models such as $f(R)$ gravity \citep{lb2007,hs2007,bbds2008} are such that the mass of the scalar field becomes large in dense bodies, effectively suppressing the magnitude of the scalar force; other models such as the dilatons \citep{bbds2010} and symmetrons \citep{hk2010,pospelov} are such that the effective coupling to matter becomes vanishingly small in dense environments. All cases in the second class of screened modified gravity can be described by the same formalism which has been recently unified \citep{bdl2011,bdlw2012}. In this paper, we will concentrate on the second class. It has been shown in \citep{bdlw2012} that the background cosmology of these models is extremely constrained. Indeed, the fact that particle masses (in the Einstein frame) and the gravitational constant (in the Jordan frame) cannot vary substantially between \g{the era of the} Big Bang Nucleosynthesis (BBN) and now implies that the scalar field must stay very close to the minimum of the effective potential since before BBN. This is guaranteed when the mass of the scalar field on the cosmological background is much heavier than the Hubble expansion rate, securing the stability of the minimum to `kicks' occurring when particles such as the electron\g{s} decouple \citep{bbdkw2004}. A consequence of this is that the effective equation of state of the scalar field in the late-time Universe becomes extremely close to $-1$, hardly distinguishable from the pure $\Lambda$-cold dark matter ($\Lambda$CDM) scenario. In practice, models of $f(R)$ gravity, chameleon, dilaton and symmetron types usually behave like $\Lambda$CDM in \HANS{the} background cosmology since before BBN. Fortunately, this does not imply that their cosmology is totally \g{degenerate with that of the $\Lambda$CDM model}: the effects of modified gravity appear \g{in the structure formation}. Indeed, within the Compton wavelength of the scalar field\footnote{\BLED{The Compton wavelength of a scalar field is defined as $\lambda\equiv m^{-1}_{\rm eff}$, and $m_{\rm eff}$ is the effective mass of the scalar field (see below).}}, gravity is modified and the growth rate of structures is altered \citep{bbdkw2004,bdlw2012}. At the linear level, this results in a modification of the growth equation which depends on the scalar field mass $m(a)$ and the coupling to matter $\beta (a)$ expressed as functions of the scale factor. It turns out that all screened modified gravity models \BLED{ with no higher derivative terms in their Lagrangian, including their \g{field-dependent} potential $V(\varphi)$ and the coupling to matter $\beta(\varphi)$}, can be {\it fully} reconstructed from the {\it sole} knowledge of the functions $m(a)$ and $\beta (a)$. This allows one to engineer models directly from their linear perturbation properties, i.e., given $m(a)$ and $\beta(a)$ one can build a fully consistent model of modified gravity defined by $\beta(\varphi)$ and $V(\varphi)$ \citep{bdl2011,bdlw2012}, which implies that one could study the nonlinear evolution of cosmic structures in the late Universe simply from the knowledge of $m(a)$ and $\beta (a)$. This provides a {\it systematic} approach to screened modified gravity which can be applied to generalised chameleon, dilaton and symmetron models. For other schemes to parameterise modified gravity see \citep{ccm2007,aks2008, jz2008,s2009,fs2010,bfsz2011}. Studying the nonlinear regime of structure formation is of particular importance for screened modified gravity models, as local gravity tests often imply that deviations from general relativity are strongest on megaparsec (Mpc) scales \citep{bdlw2012}, where nonlinearities cannot be neglected. Two competing effects influence the dynamics of modified gravity here. On the one hand, the gravitational interaction is enhanced by the presence of a long-range fifth force which implies an increase of the growth of structure. On the other hand, where local matter densities are high enough, screening effects develop and structure formation converges to its GR behaviour. These two competing effects have been confirmed in already-available $N$-body simulations of $f(R)$ gravity \citep{o2008,olh2008,sloh2009,zlk2011,zlk2011b,lh2011,lzk2012,lzlk2012,jblzk2012,lhkzjb2012}, chameleon \citep{lz2009,zmlhf2010,lz2010,li2011}, dilaton \citep{bbdls2011} and symmetron \citep{dlmw2012,wml2012} models. In this work, we apply the $(m(a), \beta(a))$ parameterisation to generalise dilaton and symmetron models and study their large-scale structure formation. We use modified versions of the {\tt ECOSMOG} code \citep{ecosmog} to run $N$-body simulations in these models. This code is based on the publicly-available adaptive mesh refinement (AMR) code {\tt RAMSES} \citep{ramses}, which is efficiently parallelised and suitable to run simulations systematically. The AMR nature of the code means that a higher resolution can be achieved, without sacrificing the overall performance of the code, in dense regions where the field equations are most nonlinear, ensuring the accuracy of the fifth force calculation there. As a result, our simulations are able to probe the structure formation in these modified gravity models down to scales well below the typical dark matter halo sizes. The results of our simulations indicate that large deviations from $\Lambda$CDM in the power spectrum can be found on scales of order 1 Mpc for both symmetron and dilaton models for values of the parameters which comply with the local constraints $($the gravitational tests in the Solar system and a mild suppression of the fifth force on galactic scales typically impose that the range of the fifth force should be less than a few Mpc in the cosmological background$)$. Large differences are also present in the number density of intermediate-sized dark matter halos with masses of order $10^{13}-10^{14}h^{-1} M_\odot$ \HANS{(representing objects from groups of galaxies to small galaxy clusters)}. For models with a fifth force whose range in the cosmological background is of order Mpc and a coupling strength to matter of order unity, the deviation from $\Lambda$CDM can reach $\sim40$\% in the symmetron case and $\sim30$\% in the dilatonic one. Such large differences are testable using future galaxy surveys. Moreover, symmetron and dilaton models are distinguishable thanks to the very different time dependence of their couplings to matter. For symmetrons, the coupling has a slow dependence on the scale factor $a$ in the recent past of the Universe and vanishes before a transition redshift $z_\ast$ (its definition will be given later). Dilaton models have a much sharper dependence on the scale factor and generically decrease exponentially fast going back in time. \BLED{As will be discussed in detail in \S~\ref{subsect:tomography}, the time dependence of the coupling strength can be roughly translated into a density dependence, and the steep density dependence in the recent past of the Universe (or equivalently in regions of low matter density) for dilaton models suggests that the dilaton screening is more efficient}. These properties make the matter power spectra and halo mass functions behave qualitatively differently in these models. \BLED{We will give a more detailed summary of the results in the concluding section.} The layout of this paper is as follows: in \S~\ref{sect:mod_grav} we review scalar-tensor theories and show how such theories of modified gravity can be analysed using a simple parametrisation which encapsulates all the dynamics; in \S~\ref{sect:models} we briefly describe the generalised symmetron (\S~\ref{subsect:symmetron_model}) and dilaton (\S~\ref{subsect:dilaton_model}) models and the possible effects of varying each model parameter; the equations that will be used in the $N$-body simulations are summarised in \S~\ref{sect:nbody_eqns}, while the details are given in \S~\ref{sect:discrete_eqns}; we next carry out tests of our codes in \S~\ref{sect:code_tests}, and the cosmological simulations of this work are then discussed in \S~\ref{sect:sim} for the symmetron (\S~\ref{subsect:sim_symmetron}) and dilaton (\S~\ref{subsect:sim_dilaton}) cases respectively; finally we summarise and conclude in \S~\ref{sect:summary}. \BLED{In the paper we use the \HANS{units $\hbar = c=1$} except where $c$ appears explicitly. Overbar (subscript $_0$) denotes the background (present-day) value of a quantity and subscript $_\varphi$ means ${\rm d}/{\rm d}\varphi$. $\kappa=8\pi G_N=M_{\rm Pl}^{-2}$, where $M_{\rm Pl}$ is the reduced Planck mass and $G_N$ is Newton's constant, are used interchangeably.} \\\\
|
\label{sect:summary} \subsection{Symmetron and dilaton screening} \label{subsect:temp} Modified gravity models vary according to their screening mechanisms by which the fifth force is suppressed in local environments. The Vainshtein mechanism works in theories of the Galileon type where a scalar field with non-canonical kinetic terms couples to matter in a reduced fashion in dense environments. Chameleons have an environment-dependent mass that becomes large enough to Yukawa suppress the fifth force in dense regions. Finally, the symmetron and the dilaton share a similar mechanism whereby the coupling of the scalar field to matter is field-dependent and can vanish in the presence of dense matter. What distinguishes these two types of models is their scalar potentials: a Mexican-hat for symmetrons and a monotonic function for dilatons. The coupling function for both types of models is a quadratic function\footnote{\BLED{Of course, other types of coupling functions can be used, as we have done in the generalised symmetron model.}}. Following the idea of \cite{bdl2011,bdlw2012}, the generalised dilaton and symmetron models studied here are completely specified by two temporal functions $m(a)$ and $\beta(a)$. These give the most general models with a quadratic coupling to matter and scalar field mass that is a power-law function of $a$ in the background cosmology for the generalised dilatons. For the generalised symmetron models, the scalar field mass vanishes for $a\leq a_\ast$ and increases to its present cosmological value from then. In both models, the screening of the fifth force is achieved in high density regions where the scalar field is trapped near the minimum of $A(\varphi)$. Yet the temporal dependences of the coupling to matter are drastically different: for generalised symmetrons it varies smoothly from a vanishing value for $a\le a_ \ast$ to its present value whereas the generalised dilatons it grows exponentially fast in the recent past of the Universe to reach its present value. As discussed in \cite{bdlw2012}, the background expansion rate of such models is practically indistinguishable from that of the standard $\Lambda$CDM paradigm, so that the cosmological effects of the fifth force could only be seen in the large-scale structures. In this work, we have performed large-scale $N$-body simulations for the generalised dilatons and symmetrons, investigating in detail the effects of varying the dilaton and symmetron parameters on the nonlinear structures of the Universe. Some of these parameters are associated with the coupling to matter $\beta_0$ (\BLED{$\beta_\star$ for the symmetron case}), and $\xi$ which specifies the range of the fifth force on the cosmological background. A few extra parameters are used in the parameterisation to define the shapes of the potential and coupling function as functions of the scalar field. For the dilatons, these parameters are $A_2,r$ and for the symmetrons they are $a_\ast$, $N$ and $M$. Let us first discuss the common features of these models: \begin{enumerate} \item The coupling to matter $\beta_0$ (\BLED{or $\beta_\star$}) determines the overall strength of the fifth forces, and increasing them leads to more structures. \item Decreasing $\xi$ leads to a shorter range for the fifth force and therefore a smaller enhancement of matter clustering\footnote{\BLED{In the dilaton case, changing $\xi$ also affects the coupling strength, making the dependence on $\xi$ more complicated.}}. \end{enumerate} \BLED{In the end, the effects on structure formation are mainly determined by how fast the fifth force evolves and how efficient it is screened in dense regions. An intuitive way to see this is to look at the expressions of $\beta(a)$ in these two models, as our discussion on tomography shows that this could be translated into $\beta(\rho_m)$, therefore giving us a sense about the screening, at least qualitatively. From Eqs.~(\ref{eq:beta_new_symmetron}, \ref{eq:beta_new_dilaton}) we can see that \begin{enumerate} \item In symmetron models, the coupling vanishes at $a\leq a_\ast$ (or equivalently for $\rho\geq\rho_\ast$) and after that it grows as a power-law function. Varying from $0$ to $\beta_\star$ between $a=a_\ast$ and today, $\beta$ depends quite sensitively on $a$ or $\rho_m$ in the regime with $\rho_m\leq\rho_\ast$; however, the symmetry of $V_{\rm eff}$ can be quickly restored for $\rho_m>\rho_\ast$ resulting in a strong suppression of the fifth force. In other words, there is a clear {\it cutoff} density beyond which the screening is very effective, and this cutoff is close to $\rho_\ast$, which is fairly low. \item In dilaton models, the coupling grows exponentially \HANS{with time and with decreasing density}. As can be seen in Eq.~(\ref{eq:beta_new_dilaton}), $\beta$ decreases and becomes vanishingly small if one goes back in time or goes to high-density regions, much more quickly than it does in the symmetron models [c.f.~Eq.~(\ref{eq:beta_new_symmetron})]. This implies that the dilaton screening can become effective for lower densities than the symmetron mechanism. \end{enumerate} It appears that the dilaton screening mechanism is more efficient than the symmetron mechanism. However, local tests of gravity are carried out in very dense regions, where the fifth force can be strongly suppressed in both models. Without specifying the exact parameter values for a given model, being it dilaton or symmetron, it is hard to say which one can satisfy local constraints more easily\footnote{\BLED{It is clear that by varying the parameter values both models can be made either more or less screened.}}.} \subsection{Summary of numerical results} Let us now summarise the results for each model. \subsubsection{Generalised symmetron models} The symmetron models we have simulated are close to what is allowed by local gravity experiments. Those constraints are mainly on the combination of the parameters $a_\ast$ and $\xi$ with the coupling strength $\BLED{\beta_\star}$ being an (almost) unconstrained parameter. This parameter, which controls the magnitude of the fifth force compared with gravity, can in principle be constrained by its effect on the cosmic structure formation. Our simulations show that for a fiducial value of $\BLED{\beta_\star} = 1.0$ the symmetron models predict an enhancement of the nonlinear power spectrum with respect to $\Lambda$CDM of up to $40\%$ for $k \sim 1~h${Mpc}$^{-1}$ and up to $50\%$ at $k\sim 10~h$Mpc$^{-1}$. Likewise we find an enhancement of up to $50\%$ in the mass function for halo masses in the range of $10^{12}-10^{14}h^{-1}M_{\odot}$. \BLED{We have shown how the fifth-force effect is changed by varying the other four model parameters: $a_\ast, N, M$ and $\xi$.} \begin{enumerate} \item \BLED{The parameter $a_\ast$ controls when the symmetry in $V_{\rm eff}(\varphi)$ is broken so the fifth force becomes non-vanishing. Decreasing $a_\ast$ gives it more time to influence the matter clustering, as a result not only the matter power spectra and mass functions deviate more from the $\Lambda$CDM results but also their shapes change qualitatively (more discussion below).} \item \BLED{$N$ is the parameter which controls the coupling strength via $\beta\propto\varphi^N$. Since $|\varphi|$ is very small, increasing $N$ will suppress the magnitude of $\beta$ (or the fifth force), and therefore causes less clustering of matter.} \item \BLED{$M$ is the {\it shape} parameter of the symmetron field potential, which determines how easy it is for $\varphi$ to roll away from $\varphi=0$ where $\beta$ vanishes. Increasing $M$ makes this easier, leading to a less-screened fifth force and thus more clustering and structures of matter.} \item \BLED{$\xi$ controls the scalar field mass and therefore the range of the fifth force in vacuum, $\lambda_\star=2998\xi h^{-1}$Mpc. Increasing $\xi$ makes the scalar field mass (range of the fifth force) proportionally larger (shorter), and thus leads to a stronger suppression of the fifth force and limits its range.} \end{enumerate} As a rough guidance, increasing the symmetry-breaking scale factor $a_\ast$ from $0.33$ to $0.50$, decreasing $\lambda_\star$ from $2.0h^{-1}$Mpc to $1.0h^{-1}$Mpc, increasing $N$ from $2$ to $4$ or reducing $M$ from $6$ to $4$ are found to lower the enhancement of the power spectra and mass functions by $\sim10-20\%$. \BLED{The parameters we adopt in the simulations are in the `realistic' range and can be tested by future galaxy surveys.} \subsubsection{Generalised dilaton models} \BLED{We have also studied how structure formation in the generalised dilaton models is affected by varying the four model parameters $A_2, \beta_0, r$ and $\xi$.} \begin{enumerate} \item The effect of increasing $A_2$ is to make the total effective dilaton potential $V_{\rm eff}(\varphi)$ steeper and so to keep the scalar field closer to $\varphi_\ast$, where $\beta$ and the fifth force vanishes. The $\Lambda$CDM limit is retrieved by letting $A_2\rightarrow\infty$. According to our simulations, reducing $A_2$ to $5\times10^4$ produces a $\sim20\%$ enhancement in the nonlinear matter power spectrum between $z=1$ and $z=0$, which is significantly smaller than the linear perturbation predictions, demonstrating the efficiency of the dilaton screening mechanism. It also enhances the mass function by maximally $\sim25\%$ in the same redshifts. These numbers assume that $\beta_0=0.5$. \item The effects of increasing $\beta_0$ are to strengthen the fifth force overall, and $\beta_0=0$ corresponds to the $\Lambda$CDM paradigm. The simulations show that even increasing $\beta_0$ to $1.0$ only causes $30-35\%$ enhancement in the matter power for scales smaller than $k\sim1h$Mpc$^{-1}$ between $z=1$ and $z=0$. This is at least $50\%$ smaller than the linear perturbation result, again showing that the fifth force is efficiently screened in dense regions. In the mean time, the mass functions are increased by up to $50\%$ with respect to the $\Lambda$CDM prediction. These numbers assume that $A_2=10^5$. \item Increasing $r$ to $3/2$ simultaneously increases the strength and decreases the range of the fifth force. The $r$-dependence of the matter clustering is rather weak as a result of the cancellation due to these two opposite effects. Assuming $A_2=10^5$ and $\beta_0=0.5$, increasing $r$ to $1.333$ only enhances the matter power spectra by less than $10\%$ at $k\sim1h$Mpc$^{-1}$ and $15\%$ at $k\sim10h$Mpc$^{-1}$, which is again significantly smaller than the predictions of linear perturbation theory. The mass function increases by up to $25\%$ in this case. \item The effects of increasing $\xi$ are similar to those of decreasing $r$, and as a result the dependence on $\xi$ is also fairly weak. \end{enumerate} \BLED{Again, future galaxy surveys can place realistic constraints on the models studied here.} \subsubsection{Highlights and comparisons} In both the generalised symmeton and dilaton models, as in $f(R)$ gravity models \citep{lhkzjb2012},we find that at late times the linear perturbation theory fails to be a good approximation even for quite large scales ($k\sim0.05h$Mpc$^{-1}$). However, at earlier times it gives better agreement with the full simulations. This indicates that the environmental suppression of the fifth force becomes more important at late times when cosmic structures (very dense matter clumps) have already formed. This highlights the importance of numerical simulations in the study of (screened) modified gravity models. \BLED{The deviations of matter power spectra and mass functions from $\Lambda$CDM in the symmetron and dilaton models are not directly comparable, because they depend on the exact parameter values used in each model. However, we can see that the shapes of $\Delta P/P$ and $\Delta n/n$ can be very different in the two models, which is probably a consequence of the different behaviour of the respective fifth forces.} \BLED{At early times, $\Delta P/P$ increases with $k$ in both models (see e.g., Figs.~\ref{fig:dpop_a_0.5_sym} and \ref{fig:dpop_a_0.3}), similarly to what we see in $f(R)$ gravity models \citep{zlk2011,lhkzjb2012}. Differences appear at late time when the fifth force has been in effect for long enough: \begin{enumerate} \item For $f(R)$ gravity models we see that $\Delta P/P$ develops a peak at $k\sim\mathcal{O}(1)h$Mpc$^{-1}$, and on even smaller scales it decreases with $k$. The peak comes from the enhanced matter clustering due to the fifth force acting between clusters, and the turnover on small scales is because (compared with $\Lambda$CDM result) on these scales the short-range fifth force still accelerates particles and prevents them from further clustering\footnote{\BLED{Contrary to intuitive understandings, this is {\it not} because `the fifth force is suppressed on small scales'. The chameleon effect only reduces the range of the fifth force, but not its amplitude within that range.}}. \item In the symmetron case, we also see the peak of $\Delta P/P$ at $k\sim\mathcal{O}(1)h$Mpc$^{-1}$, and on even smaller scales it goes up again. This seems to imply that the particle velocity inside halos stops being enhanced after the screening effect has kicked in (recall that the `cutoff' density for screening is quite low here and that `screening' here means a suppression of the amplitude, rather than range, of the fifth force), as a result of which the shape of $\Delta P/P$ on small scales is preserved since early times. \item In the dilaton models, no obvious peak of $\Delta P/P$ can be seen: the power spectrum seems to have flattened on scales smaller than $k\sim1h$Mpc$^{-1}$. Such a flattening in $\Delta P/P$ is expected in the linear perturbation results for both the symmetron and dilaton models, as in the linear regime the time at which the fifth force becomes non-negligible is scale-independent below the scale $m_{0,\star}^{-1}$. For symmetrons the flattening is destroyed by the screening effect, while for dilatons it is not. As mentioned in \S~\ref{subsect:temp}, dilaton screening can apply to lower matter densities: this indicates that the inter-cluster fifth force can be strongly suppressed as well, and thus the peak has not yet developed (notice that in some cases, such as B4, there is a small bump). Again, a more definite conclusion could only be drawn after a more detailed study of the density and velocity fields in the simulations, which is beyond the scope of this paper. \end{enumerate} } \BLED{The shape of $\Delta n/n$ at late times is similar in symmetron, dilaton and $f(R)$ gravity models, and the most important feature is that it goes down in the high-mass end, demonstrating efficient screening of the fifth force in these large structures. At early times, however, $\Delta n/n$ for the dilaton models show very weak mass dependence, which is close to the linear theory prediction, namely the fifth force is scale-independent.} \BLED{In the symmetron models, varying the parameters $a_\ast, N, M$ and $\xi$ changes the shape of $\Delta P/P$ (and of $\Delta n/n$) in similar ways, which results in a degeneracy in these parameters. This is because all these parameters control the degree of screening of the fifth force.} \BLED{This is not the case for the dilaton models, in which only a variation of $A_2$ changes the screening monotonically. Varying $\beta_0$ changes the overall strength of the fifth force more than its screening, while varying $r$ or $\xi$ changes the screening in more complicated ways. As a result there is no degeneracy in these parameters, except between $r$ and $\xi$ (see Fig.~\ref{fig:dpop_a_1.0}).} \subsection{Conclusions and outlook} \BLED{In short, the aim of this paper is threefold: \begin{enumerate} \item to show the power of the modified gravity parameterisation proposed in \citep{bdl2011,bdlw2012} in systematic studies of structure formation, \item to acquire a sense about the qualitative behaviour of the generalised symmetron and dilaton models, and the effects of varying individual parameters, and \item to make a preliminary exploration of the 4-dimensional parameter spaces in these models and find models which are testable by the near-future observations. \end{enumerate} } For all the test models in this paper, we find deviations from $\Lambda$CDM with similar magnitudes as those found in the $f(R)$ gravity model \cite{zlk2011,lhkzjb2012}, which means that many of the cosmological tests of $f(R)$ gravity \cite{lzk2012,zlk2011b,lzlk2012,jblzk2012} could in principle be carried out here as well. On the other hand, the predictions of the cosmological observables can be different from those in other modified gravity models with screening mechanisms, such as the chameleon models. \BLED{For example, the shape of the matter power spectrum can be different in the symmetron, dilaton and $f(R)$ gravity models, which implies that the respective screening mechanisms indeed work quite differently. It would be interesting to understand better the origin of such differences and see if they can be used to distinguish between the different modified gravity models in cosmology. These studies are under way.}
| 12
| 6
|
1206.3568
|
1206
|
1206.3042_arXiv.txt
|
Asteroid \tcc\ was a Near Earth Asteroid that impacted the Earth on 2008 October 7. Meteorites were produced by the break-up of \tcc\ in the high atmosphere and at present, about $600$ meteorites - called Almahata Sitta - coming from \tcc\ have been recovered. A mineralogical study of Almahata Sitta fragments shows that the asteroid \tcc\ was made of meteorites of different types (ureilites, H, L, and E chondrites). Understanding the origin of this body and how it was put together remain a challenge. Here we perform a detailed spectroscopical and dynamical investigation to show that the most likely source region of \tcc\ is in the inner Main Belt at low inclination ($i<8^\circ$). We show that asteroids with spectroscopic classes that can be associated with the different meteorite types of \as\ are present in the region of the Main Belt that includes the Nysa-Polana family and objects of the Background at low inclination. Searching for a possible scenario of formation for \tcc, we show that there is little chance that \tcc\ was formed by low velocity collisions between asteroids of different mineralogies, in the current asteroid belt. It seems more likely that the heterogeneous composition of \tcc\ was a inherited from a time when the asteroid belt was in a different dynamical state, most likely in the very early Solar System. Because ureilites are fragments of a large, thermally metamorphosed asteroid, this suggests that the phases of collisional erosion (the break-up of the ureilite parent-body) and collisional accretion (the formation of the parent body of \tcc) overlapped for some time in the primordial asteroid belt.
|
\label{section_intro} Meteorites are a partial sample of asteroids that survive the passage through the Earth's atmosphere. The identification of the source regions of the different type of meteorites is essential to be able to link the mineralogical properties of meteorites to the parent asteroids and, consequently, to address the mineralogical evolution in the asteroid belt. However, this is not an easy task and only some of these links could be established: for instance, the group of howardites, eucrites, and diogenites (HEDs) meteorites are thought to come from the Vesta family of asteroids \citep[e.g.][]{binzel93}; more speculatively, L ordinary chondrites could come from the Gefion family \citep{nesvorny09}, while asteroids of the the Flora family bear spectral similarities with the LL chondrites \citep{vernazza08}. However, the parent bodies of most meteorite types, if still intact, are unknown. The discovery and spectroscopic observation of the near-Earth asteroid (NEA) \tcc\ (henceforth \tc) 20 hours before it impacted the Earth's high atmosphere, and the subsequent recovery of meteorites (called Almahata Sitta) -- clearly coming from this body -- was a major result in this respect \citep{jenniskens09, shaddad10}. It allowed a direct link between an asteroid and meteorites to be established for the first time: the asteroid was classified as belonging to the spectroscopic F-class \citep[in the Tholen classification,][]{tholen84} or B-class \citep[in the Bus classification,][]{busbinzel02} on the basis of the flat shape of its reflectance spectrum in the region between 500 and 1000 nm. Moreover, among the 47 meteorites initially recovered, it was observed that the visible spectrum of the fragment $\#7$ matches the telescopic spectrum of \tc\ obtained before the impact with the Earth's atmosphere \citep{jenniskens09}. Fragment $\#7$ is an achondrite polymict ureilite \citep{jenniskens09}. Ureilites are in a group of achondritic meteorits that exhibit both primitive and evolved characteristics \citep{cloutis10}: in particular, they are characterized by olivine and pyroxene-rich clasts among carboneous material (mainly graphite); fine-grained graphite is also present, which lower the albedo of the meteorites \citep[about 10-12\%;][]{hiroi10}. Ureilites were initially thought to derive from $S$-class asteroids (see for instance \citealt{gaffey93}). However, because of its spectral properties, \citet{jenniskens09} propose a link with $B$-class asteroids according to the Bus classification or $F$-class in the Tholen classification. This is more plausible than a link with $S$-class asteroids, given the low albedos of ureilite meteorites, consistent with $B$ or $F$-class asteroids but not with the $S$-class. It is worth to note that the F-class can be distinguished from the B-class from a much weaker UV drop-off in the spectra of the former compared to the latter and also because B-class asteroids show a moderately higher average albedo than F-class bodies. However, in the Bus classification these two classes are merged in a unique class (the B-class). This is because the \citet{busbinzel02} spectral classification is based on spectra acquired with CCD spectrographs, which -- in general -- do not observe far enough in the UV to observe the above-mentioned drop-off feature. We will refer in this work to B-class asteroids including both Tholen B- and F-classes, and Bus B-class. Interestingly, mineralogical studies of \as\ show that, among the $\sim 600$ fragments recovered \citep{shaddad10}, about $70-80\%$ are ureilites, while the remaining $20-30\%$ are enstatite chondrites, H and L ordinary chondrites. More specifically, \citet{bischoff10} show that, among a subsample of 40 deeply studied meteorites from \as, 23 fragments are achondritic ureilites and 17 have chondritic litologies with 14 of them corresponding to enstatite chondrites, 2 to H ordinary chondrites and 1 to a new type of chondrite (see \citealt{horstmann10} for more details). Although small clasts of different types are quite common in brecciated meteorites \citep[see][for a review]{meibomclark99} and carboneous material is found in some HED meteorites, it is the first time that meteorites of very different mineralogies (i.e primitive fragments with achondrite polymict ureilites and evolved ones such as ordinary chondrites or enstatite chondrites) are associated to the same fall. This led to make the hypothesis that \tc\ was an asteroid made of blocks of different mineralogies, held together very loosely \citep[given the explosion of the body at the anomalously high altitude;][]{jenniskens09,bischoff10,shaddad10}. Tracing back \tc\ to its source region in the asteroid Main Belt would allow us to understand the origin of the Almahata Sitta meteorites and how \tc\ was put together by loosely assembling meteorites of different mineralogies. Establishing this link would also be fundamental to shed light on the source region of ureilites, that albeit rare, are the forth major class of primitive meteorites recovered on Earth after the CV, CI and CO carbonaceous chondrites. In their attempt to find the source region of \tc\ and \as, \citet{jenniskens10} selected all $B$-class asteroids, according to the Bus classification, and objects of the Tholen F and B classes and searched for spectra similarities with \tc\ and \as. As a result of their study, they showed spectral similarities between \tc\ and ungrouped asteroids as well as several dynamical asteroid groups (or families) as possible origins for the \tc\ asteroid, namely Polana ($2.4$ AU, $3^\circ$), Hoffmeister ($2.8$ AU, $4.5^\circ$), Pallas ($2.8$ AU, $33^\circ$), Themis ($3.15$ AU, $1.5^\circ$), and Theobalda ($3.2$ AU, $14^\circ$). Later, from dynamical grounds, \citet{jenniskens10} concluded that asteroids from the inner asteroid belt (i.e. with orbital semimajor axis $a<2.5$AU) are the likely parent bodies of \tc. This reduces the choice to dispersed B-class asteroids in the inner main belt and the Polana asteroid group. In Section~\ref{section_dynamics}, we revisit this issue studying the possible dynamical source regions for \tc. We recall here that the Polana group is part of a cluster of asteroids known as the Nysa-Polana family \citep{nesvorny10}, which is located in the inner main belt, between the $\nu_6$ secular resonance (at heliocentric distance $\approx$ 2.1 AU) and the 3:1 mean motion resonance with Jupiter (at heliocentric distance of 2.5 AU). This family has a complex -- twofold -- structure in orbital proper element space \citep{nesvorny10}, suggesting that it is the outcome of at least two independent break-up events in the same orbital region. From the few spectral data available at the time, \citet{cellino01} argued that the Nysa-Polana family contains asteroids of three spectral classes. The first class is that of B-class objects, like asteroid (142) Polana itself -- note that \citet{cellino01} uses the F-class classification from the \cite{tholen84} taxonomy; the second class is the S-class, with the largest member being identified as the asteroid (878) Mildred; the third class is that of X-class objects, such as the asteroid (44) Nysa. In this manuscript, we revisit this result using a much wider dataset of spectro-photometric data provided by the Moving Objects Catalog of the Sloan Digital Survey \citep[SDSS,][]{ivezic02}, which is analyzed here using a new classification algorithm \citep[][described in Section \ref{section_method}]{michel05} developed for the Gaia space mission of the European Space Agency. A detailed study of the mineralogy of the Nysa-Polana family is of great importance also for better understanding the origin of other NEOs. In partcular, \cite{campins10} claimed that the asteroid (101955) 1999 RQ$_{36}$, target of the sample return mission OSIRIS-REx (approved in the program New Frontiers of NASA), was delivered to near-Earth space via the $\nu_6$ secular resonance from the Polana group. Moreover, the binary asteroid (175706) 1996 FG$_3$, primary target of the sample return mission Marco Polo-R, under study at the European Space Mission (ESA), might have formed within the Polana group and delivered to the near-Earth space via the overlapping Jupiter 7:2 and Mars 5:9 mean motion resonances rather than the $\nu_6$ \citep[see][]{walsh12}. As a consequence, in Sections~\ref{section_polana} and \ref{background}, we perform a spectroscopic analysis using the SDSS data of the asteroids of the Nysa-Polana family as well as dispersed asteroids of the inner Main Belt (called objects of the background) in order to find spectral matches with \tc\ and \as. Finally, in Section~\ref{section_formation}, we investigate a possible formation scenario for the \tc\ asteroid as a rubble pile of rocks of different mineralogical types, which is based on the peculiar low inclination of the Nysa-Polana family and dispersed asteroids.
|
\label{section_conclusion} From our study, it seems that the Nysa-Polana family and the Background at low inclination are good candidates for the origin of \tcc\ and \as. First of all, as mentioned in Section \ref{section_dynamics}, the Nysa-Polana family is located close to the $\nu_6$ secular resonance which is the favorite route leading to primitive NEOs and more particularly to asteroid \tcc. Moreover, the proper inclination of the Nysa-Polana family is very similar to that of \tc, which should have been maintained during the transfer of \tc\ to the NEOs region through the $\nu_6$ secular resonance. We also know, from our algorithm of spectral classification (Section \ref{section_method}), that the Nysa-Polana family gathers the 3 spectral classes (S, B, and X), which are proposed analogs to \as\ fragments (Section \ref{section_polana}) under the hypothesis that ureilites are linked to B-class asteroids. More specifically, (1) the mean spectrum of B-class asteroids of the Nysa-Polana family is spectrally matched -- at least in the visible -- with available spectra of ureilitic fragments of Almahata Sitta, (2) considering space-weathering effects, the mean spectrum of S-class asteroids of the Nysa-Polana family is compatible with the spectrum of the H-chondrite fragment $\#25$, (3) a good agreement is found between X-class asteroids and enstatite chondrites from other meteorite falls (we remind that enstatite chondrites are part of \as). Of course, a comparison with enstatite chondrites from \as\ fragments would be very useful to get a definitive match between the Nysa-Polana family and \as. In Section \ref{background}, the same kind of work was performed for objects of the inner Main Belt coming from the Background at low inclination ($i<8^\circ$). We concluded that these dispersed asteroids could also be at the origin of \tc. In Section \ref{section_formation}, we tried to explain the formation of \tc\ by low velocity impacts (below $0.5$ km/s) between different mineralogies in the neighborhood of and within the Nysa-Polana family. Selecting \tc\ as the target asteroid (d=4 m), we find a probability per projectile about $10^{-17}$ impacts during its collisional lifetime (i.e. in 16.2 Myr). As a consequence, impacts at low velocity are extremely rare and there is little chance that TC3 was formed by low-velocity impacts in the current asteroid belt. This implies that the heterogeneous composition of the parent body of \tc\ has to be inherited from a time when the asteroid belt was in a different dynamical state, most likely in the very early Solar System. One could think that an asteroid of ureilite composition was contaminated by impactors of different nature when the asteroid belt was still massive and dynamically cold, so that mutual collisions were frequent and at low velocity. However, this view is probably simplistic. In fact, a body of ureilite composition needs to be formed in the interior of a large carbonaceous asteroid which underwent significant thermal alteration \citep{singletarygrove03}. This asteroid needs to have undergone a collisional disruption to expose the ureilite material in space. The same is true for the bodies of Hn composition, with n larger than 3 \citep{gopel93}. But collisional disruptions require large relative velocities, in contrast with the view of a dynamically cold belt. Thus, the asteroid belt could not be overall dynamically cold when the parent body of \tc\ formed. These considerations suggest that, conversely to what is usually thought, accretion and collisional erosion had to co-exist for some time in the asteroid belt. For this to be possible, presumably there had to be still a significant amount of gas in the system so that, although large asteroids could be on dynamically excited orbits, the orbit of their small fragments were rapidly re-circularized by gas-drag. Consequently, the mutual relative velocities of these fragments were small and a new phase of accretion was possible for them. We remark that the heterogeneity of \tc\ is not at the microscopic level; each of the meteorites delivered to the ground are of a distinct class. Thus, \tc\ seems to be an agglomeration of meteorite-sized (i.e. few dm) pebbles of different nature. Pebbles of this size are strongly coupled with the gas and are extremely sensitive to pressure gradients. They play the key role in the new models of planetesimal formations, based on the concentration of dm-size pebbles in the eddies of a turbulent disk and on the process of streaming instability \citep{johansen07,johansen09}. These models of planetesimal formation in a turbulent disk seem a priori to be particularly favorable to explain the coexistence of collisional erosion and accretion. Large planetesimals are formed by the concentration of a large number of pebbles, forming self-gravitating clumps. Once formed, the orbits of these large planetesimals are rapidly excited by the stochastic gravitational perturbations exerted by the turbulent disk \citep{ida08,morby09}. If the threshold for collisional break-up is achieved, the pebble-size fragments of these large bodies are re-injected into the game: by being concentrated into new eddies they can give origin to new large planetesimals and so forth. Admittedly, quantitative work is needed to support this scenario; also, other more classical planetesimal formation mechanism in presence of gas drag \citep{wetherillstewart93, kenyonbromley04, weidenschilling11} might explain the coexistence of erosion and accretion as well. In this respect, it will be important to understand from the observational point of view if macroscopic heterogeneity as that of \tc\ is the exception or the rule among asteroids. \tc\ is the first object of this kind that has ever been observed, but it is also the first fall of an asteroid on Earth documented live and for which an extensive and exhaustive search for meteorites has been done. So, it might not be as rare as one could be tempted to believe. Indeed, a second similar case has just been reported \citep{spurny12}. Now that the possibility for macroscopic heterogeneity is recognized, careful investigations (also conducted by remote sensing techniques) may reveal additional interesting cases. Understanding which fraction of the asteroids are of primary or secondary accretion will be a fundamental step to constrain the asteroid formation and evolution models.
| 12
| 6
|
1206.3042
|
1206
|
1206.4454_arXiv.txt
|
In 2008 {\it AGILE} and {\it Fermi} detected gamma-ray flaring activity from the unidentified EGRET source \egretsource, recently associated with a flat spectrum radio quasar (\gbsource) at z=1.762. The optical counterpart of the gamma-ray source underwent a flux enhancement of a factor 15-30 in 6 years, and of $\sim$10 in six months. We interpret this flare-up in terms of a transition from an accretion-disk dominated emission to a synchrotron-jet dominated one.\\ We analysed a Sloan Digital Sky Survey (SDSS) archival optical spectrum taken during a period of low radio and optical activity of the source. We estimated the mass of the central black hole using the width of the C\,{\sc iv} emission line. In our work, we have also investigated SDSS archival optical photometric data and UV GALEX observations to estimate the thermal-disk emission contribution of \gbsource.\\ Our analysis of the gamma-ray data taken during the flaring episodes indicates a flat gamma-ray spectrum, with an extension of up to 15 GeV, with no statistically-relevant sign of absorption from the broad line region, suggesting that the blazar-zone is located beyond the broad line region. This result is confirmed by the modeling of the broad-band spectral energy distribution (well constrained by the available multiwavelength data) of the flaring activity periods and by the accretion disk luminosity and black hole mass estimated by us using archival data.
|
Blazars are a sub-class of active galactic nuclei, emitting from radio to TeV energies. They are subdivided in two main categories: Flat Spectrum Radio Quasars (FSRQ) and BL Lacertae (BL Lac) objects. FSRQs are characterized by a flat radio spectrum in the GHz range (with spectral index $\alpha\le0.5$, where the flux density is $S_\nu\propto\nu^{-\alpha}$), and strong and broad emission lines (with equivalent width $>$5 $\AA$). BL Lac objects, on the other hand, have no or weak emission lines with equivalent width $<$5 $\AA$.\\ Blazars continuum emission originates from a relativistic jet aligned with the line of sight. Their Spectral Energy Distribution (SED) shows a double humped shape \citep{padovani1995}, with a low energy peak lying between IR and X-rays, and an high energy peak in the MeV-TeV band.\\ The low energy region of blazar spectra is associated with the synchrotron emission coming from the jet relativistic electrons. The high energy region can be modeled through inverse Compton emission (leptonic models), with seed photons coming from an external region (e.g. the accretion disk, the dusty torus), eventually reprocessed by the broad line region, or the hot corona, or from the synchrotron process itself (synchrotron self-Compton, or SSC). A detailed description of leptonic models can be found in \cite{maraschi1992,marscher1992,sikora1994}. The high energy region can also be modeled with hadronic scenarios \citep{mucke2001,mucke2003,bottcher2007}, where the very high energy protons of the jet are radiatively important. The accelerated protons produce gamma ray emission through proton synchrotron emission, the decay of neutral pions, and synchrotron emission produced by secondaries.\\ The location of the so-called ``blazar-zone'', i.e., the spatial location of the blazar SED-peaks and gamma-ray emitting region, in blazars is still a matter of debate. \cite{sikora2008} proposed that the blazar-zone is located at 3--9 pc from the central engine for the outburst of 3C~454.3 (a bright FSRQ) occurred in 2005. For the same flare, \cite{ghisellini2007} indicated, instead, a dissipation region at 0.5--0.8 pc from the central black hole (BH). From the combined study of time-dependent polarimetric radio images at 43 and 86 GHz, the optical polarimetry, and radio, optical, X-, gamma-ray light curves, \cite{jorstad2010} proposed that the low and high energy emission is located near the 43 GHz core, at a distance of the order of tens of parsecs from the BH for 3C~454.3. A similar investigation, performed for the BL Lac objects OJ287 and AO~0235+164 \citep{agudo2011a,agudo2011b}, led to similar results.\\ \cite{tavecchio2009} established that gamma-rays emitted inside the broad line region (BLR) are absorbed at E$> 20$ $\mathrm{GeV}/(1+z)$ due to the $\gamma \gamma$ interaction with the BLR photons (internal absorption). \cite{poutanen2010} refined this result, and claimed internal absorption features at E$>5$ $\mathrm{GeV}$/(1+z) and at E$> 20$ $\mathrm{GeV}/(1+z)$ in the gamma-ray spectrum of 3C~279, 3C~454.3, PKS~1510-08, and a few other FSRQ.\\ Within the leptonic scenario, \cite{ghisellini2009} show that the contribution of external photon fields, including contributions from the BLR and dusty torus to the inverse Compton emission can be parametrized as a function of the accretion disk luminosity and the dissipation distance of the emitting blob from the BH \citep{ghisellini2009}. \cite{ghisellini2010} modeled the SED of the gamma-ray loudest blazars as being emitted at 0.01--0.5 pc from the BH.\\ Using multiwavelength observations of the blazar \gbsource, we will apply the models of \cite{ghisellini2009} to further investigate the location of the blazar-zone.\\ \gbsource\ was an unidentified gamma-ray source of the Virgo region (\egretsource), detected with low significance \citep{hartman1999,casandjian2008} by the EGRET gamma-ray telescope (operating in the 30 MeV -- 30 GeV energy range, see \citealt{esposito1999}). In recent years the gamma-ray source has shown two episodes of remarkable high energy activity: at the beginning and at the end of 2008, when it was detected by the {\it AGILE} \citep{pacciani2009} and the {\it Fermi}--LAT \citep{tramatel} gamma-ray telescopes, respectively; then named \aglsource\ \citep{pittori2009}, and \fermisource\ \citep{fermicat2}. The accurate source location determined by {\it Fermi}--LAT allowed for the association of the unidentified gamma-ray source with \bzsource\, a flat spectrum radio quasar (FSRQ) included into the second edition of the Roma-BZCAT Multi-Frequency Catalog of Blazars \citep{massaro2010}. The optical counterpart of this source is \sdsssource, located at z=1.762, and the radio counterpart is named \gbsource. In the following sections, we refer to this object as \gbsource, its radio source name.\\ Here we present the results of an analysis of multifrequency data simultaneous to the {\it AGILE} campaign on the Virgo field, and to the follow-up carried out after the {\it Fermi}--LAT detection and localization. By analysing the archival data, we estimate some fundamental physical properties of the source such as the accretion disk luminosity and the supermassive black hole (SMBH) mass. In this way we can obtain a consistent picture of the source emission in the framework of leptonic models of blazars for periods of both low and high emission activity. In Sections 2 and 3 we will describe the multi-wavelength campaigns related to this source. In Section 4 we will report on the archival data. In section 5 we will present our results consisting in the determination of the BH mass, the gamma-ray light curve and spectrum, and the SED modeling.
|
We rarely have the opportunity to detect the disk emission in FSRQs, which are generally overwhelmed by synchrotron jet emission \citep[see ][ for example]{pian1999}. However, our study here suggests that we have detected accretion-disk-dominated emission in \gbsource. Granted, we cannot fully exclude the possibility that the archival SDSS and GALEX observations we have reported could be interpreted as other emission mechanisms than thermal disk emission because we lack strictly simultaneous radio observations and extended radio light curves to corroborate the assumption of low jet emission. However the assumption of disk emission remains, in our opinion, the most likely explanation for the observations.\\ The optical observations of \gbsource\ revealed an optical flux enhancement of a factor 15--30 in 6 years, signifying a shift from accretion-disk to synchrotron, jet-dominated emission. The optical spectrum obtained in the period of faint optical emission allowed the classification of the source as a FSRQ in BZCAT. We made an estimate of the SMBH mass of \gbsource\ and the accretion rate from a period of low jet activity. With these estimates, we were able to study the December 2008 flare. Modeling the observed flat gamma-ray spectrum and SED also allowed for an investigation into the location of the blazar-zone of the object.\\ As a final remark, it is worth stressing two major points. First, by definition, our estimate of R$_{diss}$ is model dependent. The location of the dissipation region was estimated assuming the parametrization proposed by \cite{ghisellini2009} for the BLR and the torus contribution to seed photons for the EC. According to this parametrization, it was possible to derive R$_{diss}$ from the luminosity ratio of synchrotron to EC emission. In fact, in the parametrization by \cite{ghisellini2009} for a dissipation region outside the BLR, the seed photons for EC fade with distance from the SMBH. We obtained two solutions. Referring to Figure~2 of \cite{ghisellini2009}, the ratio $\frac{U'_{B}}{(U'_{BLR}+U'_{IR})}$ equals the ratio of the optical to gamma-rays luminosity at the two values of the R$_{diss}$ (assuming knowledge of the magnetic energy density). For the flares of \gbsource\ reported in this article, one solution (model 2) places the blob at R$_{diss}\sim$7 pc, with only the dusty torus as the origin of seed photons. The other solution (model 1, with R$_{diss}\sim$0.2 pc) corresponds to just outside the BLR, where the contribution of seed photons from both the dusty torus and the BLR are relevant. The magnetic field is constrained in the SED modeling by the cutoff of synchrotron emission in the UV (due to the last and most energetic electrons), and by the corresponding cutoff of EC which we cannot derive directly from data (that give non-constraining upper limits at E$>$20 GeV, see Fig. \ref{fig:sed_multiepochs}). Assuming the Thomson regime, and with only one external photon field contributing to the EC, the ratio {\it f}$_{cutoff}$ between the synchrotron cutoff energy and the EC cutoff energy is proportional to $\frac{B}{\Gamma_{bulk} <\nu_{seed}>}$, where $<\nu_{seed}>$ is the typical seed photon field energy. Hence if we have constraining data at the highest energy, and with a specific geometry in the model, we can constrain B/$\Gamma_{bulk}$. However, the geometry in the model also constrains $U'_{BLR}+U'_{IR}$, hence in the ideal case we can obtain B/$\Gamma_{bulk}$ and R$_{diss}$ from the SED modeling. The ratio of synchrotron to SSC luminosity further constrains the model parameters $R_{blob}, \Gamma_{bulk}, B$, allowing one to remove the degeneracy between $\Gamma_{bulk}$ and B. In reality, we can obtain only upper limits of {\it f}$_{cutoff}$ because the data at higher energies are non-constraining. So we have only upper limits on $B/\Gamma_{bulk}$ for each model. As a consequence, the $U'_{BLR}+U'_{IR}$ could be lower than in our parametrization (we have to maintain the ratio $\frac{U'_{B} }{ U'_{BLR}+U'_{IR} }$ at the desired value). This implies that R$_{diss}$ could be higher than our evaluations.\\ The second point is that the photon field intensity is proportional to the accretion disk luminosity, and the BLR and torus location is proportional to $\sqrt L_{d} $. In all our estimations, we assume that the disk luminosity is almost steady over time, e.g., in the low state observed during the Sloan Survey in 2002, during the GALEX observation in 2007, and during the gamma-ray flares observed by {\it AGILE} at the beginning of 2008 and by {\it Fermi}--LAT at the end of 2008. This assumption could be false. In this case, we observe that the parametrization of $U'_{BLR}$ and $U'_{IR}$ reported in equation (20) by \cite{ghisellini2009} remains unchanged while varying L$_{d}$, provided that we scale the solution for $R_{diss}$ with $\sqrt{L_d}$. Therefore, variations of the disk luminosity in time, and/or systematic errors in the evaluation of disk luminosity from our SDSS+GALEX data (possibly biased by jet emission) only slightly affect our estimation of $R_{diss}$.\\ The starting point of our modeling is that the emission region is far from the SMBH (at parsec scale), and we motivate this choice with the flat gamma-ray spectrum up to energies of 15 GeV. For a different approach and results for other blazars, we refer readers to the work of \cite{tavecchio2010}. They performed a detailed study of the localization of the emission region for bright blazars making use of the variability timescale for objects showing a spectral cut-off at (10--20)/(1+z) GeV. They obtained that the variability timescale and spectra are in agreement with a dissipation region inside the BLR.\\ We note that, contrary to the model we used, some authors \citep[e.g., ][ for example]{giommi2011} model the SEDs of both FSRQ and BL Lacs with pure Synchrotron + Synchrotron Self Compton components only.\\ Model 1 ($R_{diss}\sim$ 0.2 pc), gives a variability time scale of the order of 3 days, in agreement with the flare duration estimated from the 1-day binned gamma-ray light curve. This model, however, does not properly reproduce the flat gamma-ray spectrum. In particular, in order to reproduce the observed flux at energies $>$ 10 GeV, it overestimates the spectrum for lower energies. On the contrary, model 2 ($R_{diss}\sim$7 pc) reproduces the flat-gamma-ray spectrum, but it is not clear whether the predicted variability of the order of $\sim$100 days can be associated with the duration of the gamma-ray activity period estimated from the 1-week/1-month binned gamma-ray light curves. The third model has been built by relaxing the relation $R_{blob}$=0.1$R_{diss}$ in order to preserve the variability time-scale estimated from the 1-day gamma-ray light curve (we follow the solution proposed by \citealt{tavecchio2011} for PKS~1222+216), and it still reproduces the gamma-ray spectrum. Interestingly, the size of the emitting region for PKS 1222+216 ($R_{blob}\sim5\cdot 10^{14}$cm) and for \gbsource\ differ significantly.\\ For the third model, we obtain $R_{blob}$=0.0067$R_{diss}$, in agreement within a factor of two with the prediction of \cite{bromberg2009}, that gives $R_{blob}=10^{-2.5}R_{diss}$, for the case of efficient conversion of bulk luminosity in radiation in the strong focusing scenario. With the same assumptions, \cite{bromberg2009} assume that the location of the emitting region is at $R_{diss}\sim$2.5 $(L_{jet}/10^{46}$erg s$^{-1}) (R_{BLR}/0.1$ pc$)^{-1}$ pc from the SMBH, where $L_{jet}$ is the jet power. If we invert this relation, and we make use of our result ($R_{diss}=4.8$ pc), we obtain $L_{jet}\sim 3.5\cdot 10^{46} erg s^{-1}$. We must assume that the proton--to--emitting electron ratio is of the order of 0.1 in order to reproduce such a power (in the evaluation of proton power reported in Table \ref{tab:models} we assumed one proton per emitting electron, instead). We note, however, that \cite{nalewajko2009} evaluated that efficient radiative conversion could be assumed if the product of the bulk Lorentz factor by the opening angle is $\ga$3, and according to our third model this product is 2.
| 12
| 6
|
1206.4454
|
1206
|
1206.3873_arXiv.txt
|
We have searched and reviewed all multi-wavelength data available for the region towards the gamma-ray source \fgl\ in order to constrain its possible counterpart at lower energies. As a result, only a point-like optical/infrared source with flat-spectrum radio emission is found to be consistent with all X-ray and gamma-ray error circles. Its structure is marginally resolved at radio wavelengths at the sub-arcsecond level. An extragalactic scenario appears to be the most likely interpretation for this object.
|
% The low galactic latitude gamma-ray source known as \fgl\ is one of the entries in the recent Large Area Telescope (LAT) 2-year Point Source Catalog provided by the Fermi Gamma-ray Space Telescope in the 100 MeV to 100 GeV energy range \citep{fermi}. The Fermi team points out a preliminary classification as an active galactic nucleus (AGN) of unknown type. This high-energy source is likely to have been detected as well by other observatory missions in the past in both X-rays and soft gamma-rays. The reader is referred to Table \ref{conioya} for an historical account. Attempts to find out the nature of this object based on these lower energy detections have provided no conclusive result yet. In particular, inside the \fgl\ 95\% confidence ellipse there is only one conspicuous X-ray source, namely 1RXS J205644.3+494011. Up to very recently, the identification of this ROSAT source with the luminous star BD+49 3420 was still considered plausible instead of an extragalactic origin \citep{hr09}. In addition, \citet{pmr} proposed this X-ray emitter as a possible microquasar candidate pointing out its coincidence with an intense radio source. In this work, we address the new Fermi detection together with all the multi-wavelength observational data available to try to shed light about the true origin of \fgl. \begin{figure}[tb] \includegraphics[width=\columnwidth]{fig1_NVSS.ps} \caption{Radio sources from the NVSS in the direction of the \fgl\ field of view. The angular resolution is given by the beam size of $45^{\prime\prime}$ shown at the bottom right corner. The 90\% confidence error circles of sources detected by Fermi, INTEGRAL IBIS/ISGRI (soft-gamma and hard-X rays) and Beppo-SAX are also plotted. NVSS J205642+494005 at the map centre is the only source consistent with all these error circles} % \label{fig1} \end{figure}
|
The coincidence in position within astrometric error of the X-ray source 1RXS J205644.3+494011, the flat-spectrum radio emitter NVSS J205642+494005, the point-like infrared and optical sources 2MASS J20564271+4940068 / IPHAS J205642.74+494006.7 and all high energy error boxes listed in Table \ref{conioya}, strongly points to all being the same astronomical source. This is in addition the most peculiar object within the error ellipse of \fgl. The possible connection with this gamma-ray source is reinforced after running a Monte Carlo simulation (see the methodology in \cite{romero99}) to estimate the probability of a chance coincidence between the X-ray and the Fermi sources, that turns out to be of the order of 0.4\%. On the other hand, the bright star BD+49 3420 is clearly excluded as candidate counterpart due to its non positional coincidence with most X-ray detections. The data currently available are not able to clearly find out the nature, galactic or extragalactic, of the proposed counterpart to \fgl. Nevertheless, a crude assessment is possible relying on broad band photometry. Based on X-ray spectral fits, the total hydrogen column density towards it is estimated as about $5.3 \times 10^{21}$ cm$^{-2}$ \citep{l10}. This number translates into a visual extinction of $A_V \simeq 3.9$ mag or, equivalently, a colour excess of $E(B-V) \simeq 1.3$. Taking this into account, the deredened photometric colors of the IPHAS/2MASS counterpart amount to: $r^{\prime} - i^{\prime} \geq 0.7$, $J-H\simeq -0.9$, $J-K \simeq -0.6$, and $H-K\simeq 0.4$. When checking against tables of intrinsic stellar colours (e.g. \citet{iphas,rm91}), we find that the IPHAS colour lower limit would be consistent with a main sequence star later than A2V. However, the infrared colours cannot be easily matched with any kind of normal star. Therefore, the spectrum of this source does not resemble to be stellar and the possibility of being an extragalactic object is thus reinforced. If this is the case, the point-like appearance of the optical/infrared counterpart would be typical of a distant AGN (possibly a quasar or a blazar). In this case, the observed flat radio spectrum could be naturally interpreted in terms of self-absorption of synchrotron radio photons in the compact core. There exists another outstanding radio source inside the error circle of \fgl\ as may be seen in Fig. \ref{fig1}. Its extended appearance at 20 cm wavelength suggests a possible one-sided jet structure. The 1994 VLA archival data used in this paper clearly detect the position of a compact core consistent with NVSS J205629.93+493756.4. Its precise J2000.0 coordinates are $20^h 56^m 30.25^s$ and $+49^{\circ} 37^{\prime} 59.3^{\prime\prime}$, with an estimated uncertainty of about $0.1^{\prime\prime}$ with a measured 6 cm flux density of $21 \pm 3$ mJy. In addition, the spectral index seems to be non-thermal according to \citet{specfind}. However, despite all these facts the identification of this NVSS source with \fgl\ seems much less likely than for the NVSS J205642+494005. The key reason is the absence of any hints of X-ray emission which one would normally expect as a by-product in the context of common physical scenarios with production of gamma-rays. To conclude, we have reported new radio and optical observations at sub-arcsecond scales towards the recently detected Fermi source \fgl. Together with archival data at X-ray wavelengths, we propose that the 2MASS J20564271+4940068/IPHAS J205642.74+494006.7 is the most plausible candidate counterpart of this gamma-ray source instead of the luminous star BD+49 3420, as sometimes still quoted in the literature. Further optical/infrared spectroscopic observations would be required to find the true nature of this object, although current evidence suggests that an extragalactic scenario is a likely one.
| 12
| 6
|
1206.3873
|
1206
|
1206.6777_arXiv.txt
|
We present a spectroscopic study of the incidence of AGN nuclear activity in two samples of isolated galaxies (Karachentseva, V.E. \& Varela, J.). Our results show that the incidence of non-thermal nuclear activity is about 43\% and 31\% for galaxies with emission lines and for the total sample 40\% and 27\% respectively. For the first time we have a large number of bona-fide isolated galaxies (513 objects), with statistically significant number of all morphological types. A large fraction ($\sim$70\%) of elliptical galaxies or early type spirals have an active galactic nucleus and $\sim$70\% of them are LINERs. \textbf {We find a larger fraction of AGN in early morphological types, as also found in the general population of galaxies.} Only 3\% of the AGN show the presence of broad lines (not a single one can be classified as type 1 AGN). This is a remarkable result which is at odds with the unified model even if we consider warped or clumpy tori. Finally, we interpret the large fraction of AGN in isolated galaxies as the result of \textbf {secular accretion.}
|
Along the last 25 years, many authors have studied the AGN environment \citep{1982ApJS...50..517S, 1982ApJ...262...66S, 1984PhDT.........5D, 1985ApJS...57..643D, 1984ApJ...279L...5K, 1995IAUS..164..434D, 1996A&A...308..387D, La95, Dul99, Kron01, Kron02, 2001AJ....122.2243S, 2006A&A...451..809S, 2006ApJ...651...93K, 2006ApJ...639...37K, 2008RMxAC..32..150M}. The incidence of nuclear activity in galaxies and their environment has become a topic of debate because there are different mechanisms that can possibly trigger nuclear activity depending on the galaxy's environment. Interactions between galaxies are well known to produce enhancement in star formation in galaxies \citep{1984ApJ...287...95L, 1987AJ.....93.1011K, 1993AJ....106.1771K,Kron02, 2000ApJ...530..660B, 2007AJ....134..527W, 2007ApJ...660L..51L}. Others authors also have evidence for a connection between circumnuclear starburst and AGN \citep[and references therein]{2008RMxAC..32..139S}. There are also suggestions for a connection between interactions and nuclear no-thermal activity specifically of type 2 \citep{Dul99,Kron01}. Other studies have dealt with the AGN population of compact groups \citep{Mar10} as well as in larger groups. The purpose of this project is to investigate the conditions to trigger AGN activity in different environments. In this paper we study the incidence of activity in isolated galaxies. In a forthcoming paper \citep{HI13} we will present the results of a survey of AGN in paired galaxies of similar mass. It is important to study galaxies in a restricted environment in order to elucidate what mechanisms could be determinant to trigger AGN activity. Isolated galaxies can be defined as those systems that are formed in low galactic density environments, but that evolved without major interactions with other galaxies of not only similar mass over the last 3 Gyr. In this context, any non-axisymmetric structures in these galaxies such as bars, tails, plumes or stripping material must be the result of secular evolution.\\ The study of truly isolated galaxies is thus fundamental to benchmark the role of interactions in nuclear activity. Studies of field galaxies (e.g. Ho et al. 1997) cannot provide this information as these samples may include galaxies that have undergone or are undergoing an interaction. \textbf {This is the first paper of a series involving a self consistent and homogeneous way to study nuclear activity in galaxies in different environments}. In the present work, we study the incidence of nuclear activity in two samples of bona-fide isolated galaxies using an efficient way to extract the stellar contribution (host spectrum). The purpose of this work is to have a well defined sample of isolated galaxies with optical spectroscopic characteristics that allow us to classify them according to their type of activity. \textbf {Studying the incidence of the nuclear activity in isolated galaxies alone is of great value to establish if AGN is a common and/or persistent phenomenon even when strong tidal external perturbations have not been present during the last few Gyr of galaxy evolution. This would indicate that AGN activity can be triggered by secular evolution processes in galaxies. Our results on this sample will be further used as a benchmark to compare the incidence of activity in a sample of isolated pair of galaxies \citep{HI13}. Our samples of isolated and paired galaxies have been chosen to have consistent properties with each other except for the presence of a companion.} This paper is organized as follows. In \S 2 we describe our samples. In \S 3 we present the data analysis and classification. Results are given in \S 4 and \S 5 contains the discussion about the possible mechanisms for developing an AGN in isolated galaxies.
|
The large number of galaxies of all morphological types permits us to quantify the link between morphology and nuclear activity. Our results indicate a close link between these two properties. This implies that any result of the incidence of activity without this consideration reflects the particular morphological distribution of the sample and therefore is not reliable. Our sample includes for the first time a statistically significant number of isolated early type galaxies (E+S0). \textbf {We found that Elliptical and SO galaxies have the highest incidence of nuclear activity in isolated environments when only galaxies with emission lines are considered (a similar result was found by \cite{Var04, Co11, Sa12}). However, when the total sample is taken into account (including galaxies without emission lines) these apparent excess disappears and all early types (including Sa and Sb types) have similar fractions (see Fig. 4). This important difference could be found thanks to the large number of elliptical and spheroidal galaxies in our sample. This results are consistent with those found for ``field" galaxies \citep {1980A&A....87..142H, 1983ApJ...269..466K, Ka03, 2003ApJ...597..142M}. Finding the same trend between isolated and field galaxies could be expected if we consider that AGN require SMBHs and black holes are correlated with bulges. However, the large number of AGN among isolated galaxies is of great importance and indicates that secular evolution processes can trigger/maintain low luminosity AGN activity (see below).} We consider as secular evolution the following mechanisms capable of driving gas into the nuclear region: 1) Minor mergers (luminosity ratio larger than 10 in our sample); 2) Dark matter accretion; 3)Non-axisymmetric gravitational perturbations. With respect to dark matter, \citet{2010MNRAS.404L...6H} showed from an analytical treatment of the accretion rate that, for the largest black hole masses of quasi-stellar objects ($>$10$^{9}$ M$_{\odot}$), the runaway regime would be reached on time-scales wich are shorter than the lifetimes of the halos in question for central dark matter densities in excess of 250 M$_{\odot}$/pc$^{3}$. This limit scales inversely with the black hole (BH) mass. \textbf {The most common non-axisymmetric internal potential is due to the presence of a bar. However in the particular case of barred galaxies it has been shown that most probably bars do not enhance nuclear activity \citep{1995ApJ...438..604M, 1997ApJ...487..591H, 2012arXiv1203.1693L}. However the samples used in those studies were not rigorously isolated and thus the effects of the environment cannot be disentangled from those of the bar. The samples used in this work provide the opportunity to perform a rigorous test of the effect of a bar. This can be achieved due to both the selection criteria and the quality of the data. A detailed analysis of the bar fraction requires a deep photometric study. This analysis will be presented in a forthcoming paper \citep{Her-Tol13}.} \textbf {We note that although a large fraction of isolated galaxies are active, their SMBH has not grown significantly over the last 3 Gyr. Given that most of our galaxies are representative of the low luminosity end of AGN, the mass accretion rate should be in the range of $10^{-5}-10^{-3}$ M$_{\odot}$/yr, and the radiative efficiency $\eta$ should be significantly smaller than 10 \% \citep{2003ASPC..290..379H, 2009ApJ...699..626H}. Such low efficiencies are predicted for low luminosity AGN \citep{2008NewAR..51..733N}. Assuming the AGN in our sample have accreted at a constant rate over the last 3 Gyr the growth of their SMBH ranges between $~10^{4} - 10^{6}$ M$_{\odot}$. Then, it is clear that isolated galaxies in poor environments have failed to accrete enough material (at least during the last Gyr) to present higher luminosities and significant black hole growth. Thus, our results support a hierarchical scenario in which the environment is crucial to determine properties such as luminosity, mass, and central SMBH mass, fulfilling the expectations of the downsizing for SMBH growth \citep {2010hsa5.conf..337P}.} The spiral isolated galaxies will not migrate from the blue to the red sequence since feedback is not efficient in these faint AGN \citep{2007ApJ...659.1022K}. The later result supports again that secular evolution in these galaxies is the important mechanism to establish the bulge-black hole relation. This is in contrast to the case of massive galaxies transitioning from the blue to the red branch of the color-color diagram which require a major merger followed by a substantial feedback in the QSO phase. On the other hand our isolated ellipticals are already in the red-branch of galaxies that probably have experienced a major merger in the distant past. The fact that essentially all of them are AGN may simply reflect the fact that it is easier to drive gas to the center of spheroidal systems. The remanent inter stellar medium (ISM) in these galaxies is typically in the range (10$^{6}$-10$^{7}$ M$_{\odot}$). Therefore, they contain enough gas to power their SMBH over the last 3 Gyr. This does not exclude, however, the possibility of an external supply of material as suggested by several authors \citep{1992ApJ...401L..79B, 2000ApJS..127...39C, 2006MNRAS.366.1151S}. The large fraction of AGN in these galaxies suggests that the presence of a large bulge facilitates the mass in fall to the center. We note, however, that a small fraction of this LINERs could be fake AGN (``retired galaxies'', \citet{2011MNRAS.413.1687C} ). Finally the absence of type 1 AGN in these samples of isolated galaxies is remarkable and at odds with a simple interpretation of the unified model (UM). There is not a single type 1.0 AGN among the 175 active galaxies in our samples. The fraction of types 1.5-1.9 is less than 3 \% in both isolated galaxy samples. \textbf {All these results indicate that the presence of AGN activity is a common phenomenon in galaxies independently of the environment. This is an important result as traditionally it has been assumed that an external perturbation is required to induce nuclear activity. Our results indicate that a low luminosity AGN phase is a part of the secular evolution of a large number of galaxies. These findings are consistent with those by \citet[and references therein] {1997ApJS..112..315H, 1997ApJ...487..568H, 2002ASPC..284...13H}. However, in those studies it was impossible to disentangle the environmental effects from those of internal galactic evolution, given that the isolation history was not known a priory in their samples. Our results do not deny the possibility that external perturbations may enhance the frequency of nuclear activity among galaxies, as has been suggested by previous studies \citep{Dul99, Kron02, Kron03, Ro09, 2010MNRAS.407.1514E, 2011MNRAS.418.2043E}. The effect of a strong gravitational interaction will be studied in a forthcoming paper where the incidence of AGN in a sample of isolated close pairs of similar mass galaxies will be analyzed. We also note that the lack of high luminosity AGN in our samples points towards a dependence between environment/interactions and AGN luminosity. In this scenario, the extremely low fraction of type 1 AGN can be understood if a BLR (broad line region) can be formed only at higher accretion rates/luminosities \citep{2000ApJ...530L..65N,2009ApJ...701L..91E}. The appearance of a type 1 nucleus may be delayed by as much as 1 Gyr, as required by the evolutionary model proposed by \citet{2007ApJ...659.1022K}, where an interaction triggers first a circumnuclear starburst, and subsequently non-thermal nuclear activity. For the brightest end of nuclear activity a similar evolutionary trend is possible (from ULIRGs to luminous quasars). In this case a major merger would be required, affecting the overall properties of the host galaxy and moving it to the blue branch of the color-color diagram.} \textbf {If AGN in isolated galaxies have low accretion rates, low efficiencies, low luminosities and almost a complete absence of broad lines in their spectra, it is probable that the BLR under these circumstances is not even able to form. This is in accordance with the result by \citet {2003ApJ...583..632T, 2003ASPC..290...31T} for the absence of broad components in polarized light for $\sim$50$\%$ of the galaxies in his sample. Several studies show evidence that the Sy2s with and without broad lines in polarized lines (in other words with and without a hidden BLR) are truly different in other respects as well \citep [e.g.][]{2001RMxAA..37....3G, 2001A&A...366..765G, 2012arXiv1209.0274B}.}
| 12
| 6
|
1206.6777
|
1206
|
1206.1157_arXiv.txt
|
{NGC 6809 is a luminous metal-poor halo globular cluster that is relatively easy to study due to its proximity and low concentration. Because of its high Galactic latitude ($b = -23^\circ$), interstellar reddening and contamination is not very high. } { We aim to determine the relative proper motion and membership probability of the stars in the wide area of globular cluster NGC 6809. To target cluster members reliably during spectroscopic surveys and both spatial and radial distributions in the cluster outskirts without including field stars, a good proper motion and membership probability catalogue of NGC 6809 is required.} {The archival data of two epochs with a time-base line of 7.1 years have been collected with Wide Field Imager (WFI) mounted on the 2.2m MPG/ESO telescope. The CCD images of both epochs have been reduced using the astrometric techniques as described in Anderson et al. (2006). The calibrated $UBVI$ magnitudes have been derived using Stetson's secondary standard stars.} {We derived the relative proper motion and membership probabilities for $\sim$ 12600 stars in the field of globular cluster NGC 6809. The measurement error in proper motions for the stars of $V \sim 17$ mag is 2.0 mas~yr$^{-1}$, gradually increasing up to $\sim$3 mas~yr$^{-1}$ at $V=20$ mag. We also provide the membership probability for the published different types of sources in NGC 6809. An electronic catalogue \thanks{Full Table 5 is only available in electronic form at the CDS via anonymous ftp to cdsarc.u-strasbg.fr(130.79.128.5) or via http://cdsweb.u-strasbg.fr/cgi-bin/qcat?J/A+A/} with proper motion and membership probability for the stars will be available to the astronomical community.} {}
|
\label{INTR} Globular clusters have long been used to study the structure and formation of our Galaxy. NGC~6809 (M~55) is a sparse, metal-poor globular cluster whose proximity ($\sim$ 5 kpc) makes it an excellent target for an in-depth study of its stellar population. NGC~6809 harbours several interesting objects. Bassa et al.\ (2008) found 16 $X$-ray sources within the half-mass radius (2\arcmin.89) of NGC~6809, of which eight or nine are expected to be background sources. On the basis of optical counterparts, these authors identified three sources related to the cluster. The brightest $X$-ray source of this cluster is classified as a dwarf nova. Blue stragglers in NGC 6809 have been studied by Lanzoni et al. (2007) and exhibit a bimodal radial distribution with a central peak, a broad minimum at intermediate radii, and an upturn outwards. A detailed study about the evolved stars, including asymptotic giant branch, horizontal branch, and upper red giant branch (RGB) stars was presented by Vargas et al. (2007). A proper-motion (PM) study of NGC~6809 was conducted by Dinescu et al. (1999) using photographic plates. The authors determined an absolute proper motion ($\mu_{\alpha}cos\delta=-1.57\pm0.62$ mas yr$^{-1}$, $\mu_{\delta}=-10.14\pm0.64$ mas yr$^{-1}$) using a sample of $\sim$ 600 cluster stars brighter than $V\sim16$ photographic magnitude and background galaxies as a reference. Recently, a proper motion study of NGC~6809 was performed by Zloczewski et al. (2011) (hereafter, Zl11). They determined membership probabilities for 16,645 stars in the central part of the cluster ($8\farcm83$ $\times$ $8\farcm83$) and found an absolute proper motion $\mu_{\alpha}cos\delta=-3.31\pm0.10$ mas yr$^{-1}$ and $\mu_{\delta}=-9.14\pm0.15$ mas yr$^{-1}$. Despite the extensive photometric studies of this cluster, there is a lack of studies that provide proper motions and membership probabilities in the wide field region for NGC 6809. Wide-field images allow us to map completely any open or globular cluster in our Galaxy and its tidal trails and allow us to obtain accurate photometry for an enormous number of stars. In combination in photometric data, membership information is very useful to keep the field star contamination to a minimum. The archival wide-field multi-epoch observations taken with the WFI@{\bf 2.2}m telescope offer new opportunities to derive precise PMs with only a few years of time span, deeper by several magnitudes than previous photographic surveys (Anderson et al. 2006; Yadav et al. 2008; Bellini et al. 2009). The main purpose of the present study is to determine accurate relative PMs and membership probabilities for stars brighter than $V\sim20$ mag in the wide area of NGC 6809. Membership probabilities of different sources in NGC 6809 are also discussed. Fundamental parameters of the cluster taken from Harris (1996) are listed in Table~\ref{par}. The PMs, membership probabilities, and photometric $U,B,V$ and $I$ magnitudes are provided to the astronomical community for follow-up studies. Our membership-probability catalogue is in wide field ($26\arcmin \times 22\arcmin$) and contains $UBVI$ magnitudes, while Zl11 have provided membership probabilities only in $V$ magnitude and in the central area of $8\farcm83$ $\times$ $8\farcm83$. The structure of the article is as follows. Data taken for the present study, their reduction procedures and comparison of photometric and astrometric data are described in Sect.~\ref{OBS}, where we also determine PMs and differential chromatic refraction. In Sect.~\ref{MP} we present the cluster membership analysis. In Sect.~\ref{app} we use our catalogue to confirm the membership of previously found variables, blue stragglers, and X-ray sources. Finally, Sect.~\ref{catl} describes the catalogue while Sect.~\ref{con} represents the conclusions of the present study. \begin{table} \caption{Fundamental parameters of NGC~6809 taken from Harris (1996). } \centering \label{par} \begin{tabular}{cc} \hline\hline Parameters & Values \\ \hline $\alpha$(J2000) & 19$^{\rm h}$ 39$^{\rm m}$ 59.$^{\rm s}$4 \\ $\delta$(J2000) & $-30^\circ$ 57$\arcmin$ 44$\arcsec$ \\ $l$ & $8\fdg80$ \\ $b$ & $-23\fdg27$ \\ $\rm [Fe/H]$ & $-1.81 $ \\ $E(B-V) $ & 0.08 \\ $(m-M)$ & 13.87 \\ \hline \end{tabular} \end{table}
|
\label{con} We provided a catalogue of precise proper motions and membership probability of stars in the wide-field region of globular cluster NGC~6809. We have obtained precise proper motions and astrometric membership probabilities down to $V\sim20$ mag in 26$\times$22 arcmin$^2$ area around the globular cluster NGC 6809. Finally, we provided the membership probability for different types of variable stars, blue stragglers, and $X$-ray sources. We also demonstrated that the CCD observations taken just seven years apart can provide accurate proper motions. These proper motions are used to separate the cluster members from field stars down to$V\sim20$ mag.
| 12
| 6
|
1206.1157
|
1206
|
1206.0713.txt
|
\noindent Cosmic rays are the highest energy particles available for our study and as such serve as excellent probes of the effects of Lorentz Invariance Violations, which are expected to increase with energy. This general paradigm is investigated in this paper by studying the effects of such violations within the Coleman-Glashow model in which each particle species may have its own maximum attainable velocity, even exceeding that of light \textit{in vacuo}. The particular focus here is that the muon neutrino may have the maximum speed exceeding that of light. We show that such an assumption leads to the elongation of the decay lifetime of the pion that increases with energy over and above the time dilation effects. We provide a transparent analytical derivation of the spectral intensities of muon neutrinos and muons generated in the Earth's atmosphere by cosmic rays. In this derivation we not only account for elongation of the pion lifetime, but also for the loss of energy by the neutrinos by radiation of the electron-positron pairs through the Cohen-Glashow process, during their propagation. We then compare the theoretical spectra with observations of neutrinos and muons from large instruments like IceCube and BUST to set a limit of $\sim10^{-13}$ on the fractional excess speed of neutrinos over that of light. We also show that the ratio of the spectral intensities of downward and upward moving neutrinos at various angles constitute a diagnostic exclusively for the Cohen-Glashow process, which may be searched for in the IceCube data set. We conclude the paper with several comments, including those related to improvements of these tests when definite signals of GZK neutrinos will be observed.
|
\noindent The study of several exciting aspects of high energy astrophysics and indeed of many subtle aspects of basic physics have been given a boost by the commissioning of large detectors of cosmic ray secondaries such as ANITA, IceCube, Auger, BUST, Kolar Gold Fields, Kamiokande and other experiments with collecting powers of $\sim100$ km$^3$ \cite{Krishnaswamy1982, Krishnaswamy1975, Krishnaswamy1971, Gorham2010, Hoover2010, Becker 2011, Aglietta2000, Gray2011, Aharmim2009, Abbasi2011, Novoseltsev2010, Abraham2008, Abbasi2004, MultiKm3, Fechner2009}. These detectors have already detected $\sim10^9$ cosmic-ray muons of median energy $\sim2\times10^4$ GeV and $\sim10^4$ neutrinos that allow the spectra to be determined up to $\sim10^6$ GeV. The physics input regarding high energy nuclear interactions from accelerators, colliders and other sources help in reliably modeling the propagation of cosmic-rays through the atmosphere and qualitatively account for the observed spectral intensities of the muons and the neutrinos. Comparison of these spectral intensities with the model predictions then allow us to probe into primary cosmic ray composition at high energies and search for effects due to new physics, such as small violations of the Lorentz Invariance that may manifest themselves only at the highest energies. This paper is devoted to such an exercise. Violation of Lorentz Invariance is studied from two distinct perspectives. The first is exemplified by Michaelson-Morley and Hughes-Drever experiments which test the existence of preferred frames of reference and the anisotropy of Machian type long-range interactions of matter in the laboratory with astronomically distant matter. The extraordinary accuracy achieved in such interactions validated relativistic theories of gravity, especially GR \cite{Will1981,Will2006}. The other perspective is exemplified by the theoretical considerations of Coleman and Glashow \cite{Coleman1997, Coleman1999}, who accept the possible existence of a preferred frame, such as the frame in which the dipole anisotropy of the universal microwave background at 2.7K vanishes. In the preferred frame, the laws of physics are assumed to be invariant under translations and rotations. However they investigate the possibility that different particles could have maximum attainable speeds different from that of light \textit{in vacuo}, and these speeds could, in principle, exceed that of light by a small amount. Coleman and Glashow have developed a perturbative framework to discuss the violations of Lorentz Invariance (LIV) with terms that are renormalizeable and are invariant under the $\textup{SU} (3)\times\textup{SU}(2)\times\textup{U}(1)$ gauge symmetry of the standard model. Going beyond the standard model, Kostelecky and collaborators have carried out extensive analysis of models where Planck scale physics yields signals in the propagation of photons, neutrinos and other particles that have the potential for being observed in present-day or future experiments \cite{Kostelecky2011a, Kostelecky1989, Kostelecky2011b}. These later papers provide a comprehensive overview of the physics and the observational status of these models: Observations of high-energy gamma rays from distant astronomical sources have also been used to set lower bounds on the energy scale at which quantum gravity effects lead to increase in the velocity of light with energy \cite{Biller1999}. The aim of this paper is to discuss the bounds on LIV derived from cosmic-ray observations based on the formalism developed by Coleman and Glashow \cite{Coleman1997,Coleman1999}. In this context, we may refer to the elegant review of earlier work by Bietenholz \cite{Bietenholz2011}. We begin by recalling briefly the earlier efforts in the field of cosmic rays to search for the effects of superluminal velocities. An excellent review of the efforts to observe tachyons \cite{Bilaniuk1962} in cosmic ray showers is provided by R. W. Clay \cite{Clay1962}. The air-shower group of the Tata Institute of Fundamental Research pioneered these studies by searching for energetic particles that arrive at the air-shower array some 10-50 $\mu$s before the main shower front of electron-positron pairs, muons and gamma rays initiated by cosmic ray particles of $\gtrsim10^6$ GeV \cite{Ramana1971, Wolfendale1973}. Following the lead given by Coleman and Glashow \cite{Coleman1997}, with specific reference to the present paper, the early bounds on LIV using horizontal air showers were obtained by Cowsik and Sreekantan \cite{Cowsik1999}; detailed comments on this paper may be found in the papers of Coleman and Glashow \cite{Coleman1997, Coleman1999} and of Halperin and Kim \cite{Halperin1999}. This later paper maps the violations of Lorentz invariance into violations of the Equivalence Principle. In a subsequent paper, Cowsik et. al. \cite{Cowsik1999b} have investigated the possibility that if similar effects can induce $\nu_\mu\rightarrow \nu_e+\gamma$, then such a rate is far more strictly bounded. Stecker and Glashow \cite{Stecker2001} discuss the bounds on LIV of electrons based on observations of energetic cosmic rays. Similarly Stecker and Scully \cite{Stecker2005, Stecker2009} have put bounds on LIV in the hadronic sector by consideration of the GZK cut-off \cite{Griesen1966, Zatsepin1966}. Direct observations of the neutrinos from supernova 1987A \cite{Hirata1987, Bionta1987} allowed Stodolsky \cite{Stodolsky1988} and Longo \cite{Longo1987} to set bounds on any excess speed of neutrinos over that of light at the $\sim10^{-8}$ level, a significant improvement over early results at accelerators \cite{Adam2011}. In the context of OPERA experiments \cite{Alexandre2012,Adamson2007, Adam2011, Cacciapaglia2011, Haridass2011, Li2011, Gilles2011, Drago2011, Alexandre2011, Bi2011} several ideas of interest have been put forward and we reference a few for completeness \cite{Ellis2008, Adam2011, Giudice2012, Hagen1987, Grossman2005, Alfaro2005, Cowsik1999, Kostelecky2011a, Hambye1998a, Hambye1998b, Kostelecky1998, Kostelecky1999, Bluhm2000, Pas2005, Hollenberg2009, Coleman1998, Stecker2005, Maccione2009}. The particular focus here is to provide an analytical calculation of the spectral intensity of muons and muon neutrinos arising from the decay of pions produced by cosmic rays in the Earth's atmosphere. In carrying out these calculations, we have included the enhancement of pion life-time and the decrease in the average energy transferred to the neutrino in pion decay due to any posited superluminal motion of the muon neutrino. Secondly, we have included the effect of such a neutrino losing energy by emitting electron-positron pairs during its flight even through vacuum, as pointed out recently by Cohen and Glashow \cite{Cohen2011} in the context of OPERA experiments. We will not embark here on the ambitious program of getting the best limits for the rich variety of the LIV modifications for the various particles involved. Instead what we will attempt here is more limited and yet clearly illustrates the potential reach of this approach. Here we will focus on the effects of modifying the energy-momentum relation for $\nu_\mu$ only to $ E_{\nu} = p_{\nu} (1+\alpha )$, as suggested by Coleman and Glashow for modeling violations of Lorentz Invariance. Our analysis presented here exclude values of $\alpha$ values down to $\sim10^{-13}$. This is achieved by providing a transparent analytical calculation for the propagation of cosmic rays in the Earth's atmosphere that accurately reproduces the known data when no anomaly is assumed, and then comparing the theoretical spectra for various values of $\alpha$ with the observational data.
|
\noindent In this paper we have developed a transparent analytical model for the propagation of cosmic rays in the Earth's atmosphere that explicitly includes the effects of a superluminal motion of the muon neutrino on the decay probability and kinematics of pion decay and energy losses suffered by the neutrino through the Cohen-Glashow process as it propagates from the production site in the atmosphere onwards to the proximity to the detectors placed deep underground. The available observational data to date on the cosmic ray generated neutrinos and muons place a bound of $\alpha\lesssim10^{-13}$. We have pointed out how detectors like IceCube may search exclusively for signatures of the Cohen-Glashow process by observing the forward-backward ratio of high energy neutrinos arriving at the same zenith angles but from upper and lower hemispheres. Keeping in mind that the data published by the IceCube collaboration is limited to those acquired in 2009 and earlier years, their full data up to the present date should be able to extend the spectra to $\gtrsim10^6$ GeV and the bounds on $\alpha$ to $\sim10^{-14}$. The observations of GZK neutrinos with $E \sim 10^9$ GeV will push $\alpha$ to well below $\sim10^{-23}$!! %
| 12
| 6
|
1206.0713
|
1206
|
1206.3222_arXiv.txt
|
Quasi-static speckles are a current limitation to faint companion imaging of bright stars. Here we show through simulation and theory that an adaptive pupil mask can be used to reduce these speckles and increase the visibility of faint companions. This is achieved by placing an adaptive mask in the conjugate pupil plane of the telescope. The mask consists of a number of independently controllable elements which can either allow the light in the subaperture to pass or block it. This actively changes the shape of the telescope pupil and hence the diffraction pattern in the focal plane. By randomly blocking subapertures we force the quasi-static speckles to become dynamic. The long exposure PSF is then smooth, absent of quasi-static speckles. However, as the PSF will now contain a larger halo due to the blocking, the signal to noise ratio ({\it SNR}) is reduced requiring longer exposure times to detect the companion. For example, in the specific case of a faint companion at $5\lambda/D$ the exposure time to achieve the same {\it SNR} will be increased by a factor of 1.35. In addition, we show that the visibility of companions can be greatly enhanced in comparison to long-exposures, when the dark speckle method is applied to short exposure images taken with the adaptive pupil mask. We show that the contrast ratio between PSF peak and the halo is then increased by a factor of approximately 100 (5 magnitudes), and we detect companions 11 magnitudes fainter than the star at 5$\lambda/D$ and up to 18 magnitudes fainter at 22.5$\lambda/D$.
|
Detecting the faint reflected or self-luminous signal from extrasolar planetary companions close to a bright parent star is a technically difficult task. With the development of sophisticated image analysis and Adaptive Optics (AO) systems on several modern large (8~m class) telescopes it is now possible. AO is required to both increase the peak intensity of the point-spread function (PSF) and to concentrate the photons which are scattered into a diffuse halo by the atmosphere back into the diffraction limited core. Dedicated high-contrast imaging instruments, such as HiCAIO (Subaru, \citeauthor{Hodapp08}, 2008), SPHERE (VLT, \citeauthor{Beuzit08}, 2008) and GPI (Gemini, \citeauthor{Macintosh06}, 2006), are designed to incorporate eXtreme AO (XAO) systems and sophisticated coronagraphs to reject the light from the star whilst conserving the few photons from the angularly separated companion. Quasi-static speckles, mimicking the signal of faint companions, are now limiting the detection capabilities of these instruments (e.g. \citeauthor{Fitzgerald06}, 2006; \citeauthor{Soummer07}, 2007). Quasi-static speckles in the focal plane are caused by non-common path errors and uncorrected aberrations in the primary mirror and other optical and mechanical components. If these aberrations were entirely deterministic they could be subtracted. However, quasi-static speckles are slowly varying aberrations making calibration difficult. The current most popular methods for quasi-static speckle reduction are PSF subtraction techniques such as, PSF estimation (e.g. \citeauthor{Lafreniere07}, 2007) angular differential imaging (ADI, \citeauthor{Marois06}, 2006), simultaneous spectral differential imaging (SSDI, \citeauthor{Smith87}, 1987) and, in the case of reflected light, polarimetric differential imaging \citep{Seager00}. SSDI and PDI both require certain properties from the target. ADI is more generic but if the speckles evolve during the observation the suppression provided by the technique reduces dramatically. The temporal decorrelation timescale of these quasi-static speckles is an important factor when estimating the performance of image subtraction techniques which have proved themselves to be efficient at Strehl ratios of the order of 20--40\% \citep{Martinez12}. At high Strehl ($>$80\%) the quasi-static speckles will become even more dominant as the PSF halo is reduced further. In this high Strehl regime the speckle coherence timescale is unknown. \citeauthor{Martinez12} show that the quasi-static speckles become unstable over a timescale of a few seconds on a laboratory XAO test bench. It is thought that the evolution of the speckle pattern was primarily caused by temperature fluctuations and so, on a dedicated instrument, this could be more controlled. Several other interesting and inventive techniques are also being developed in order to further enhance the probability of detecting faint companions. \cite{Ribak08} demonstrate it is possible to enhance the contrast by placing a rotating eccentric mask in the pupil plane. This breaks the symmetry between the telescope pupil and the focal plane causing the quasi-static speckles to move, the companion will however stay fixed at the same location. \cite{Gladysz08,Gladysz09} use the statistical distribution difference of on-axis and off-axis PSF to differentiate between real sources and speckles. It is likely that only the combination of several of these techniques in conjunction with AO and a coronagraph will result in the highest contrast ratios in modern instrumentation. Here we propose to combine the idea of breaking the symmetry of the optical system and the sensor by changing the pupil function (similar to \citeauthor{Ribak08}) with an adaptive pupil mask (APM) \citep{Osborn09}. The APM is positioned in the conjugate plane of the telescope pupil. This pupil mask consists of a number of independently controllable elements. The simplest design would be a segmented mirror where each segment can either reflect the light on-axis into the remaining optical system or off-axis into a baffle. This is also similar to the speckle decorrelation (or phase--boiling) method \citep{Saha02}. The phase--boiling method involves adding additional phase aberrations to the optical path in order to force the quasi-static speckles to be more dynamic. This was later disproven by \cite{Sivaramakrishnan02} as the original phase aberrations are still present and so the quasi-static speckles are also still present, simply hidden within a field of dynamic speckles. Here we boil the speckles by introducing amplitude aberrations instead of adding phase aberrations, an important difference as the interference pattern will no longer contain the same speckle pattern, but will actually be completely different. By changing the shape of the pupil we modify the diffraction pattern in the pupil plane. Light that once interfered constructively to form a quasi-static speckle at a given location will now not. However, speckles will be formed in other locations in the focal plane. The quasi-static speckles will be forced to be dynamic, removing any dependance on speckle timescale. By changing the pupil function many times during an exposure these speckles will average out into a smooth PSF, albeit with an additional halo due to diffraction through the pupil mask (Babinet's principle). This will mean that longer integration times are required for a companion to become visible with the same signal to noise ratio (SNR). However, the quasi-static speckles will be substantially reduced, reducing the complexity of the companion identification problem. In addition to the APM reducing the static speckles into a smooth halo, we can also expose after each configuration of the mask separately. As the quasi-static speckles are now dynamic we can use other image manipulation methods to further enhance the image. Dark speckle (DS) imaging, first proposed by \cite{Labeyrie95}, is a technique designed for the detection of faint companions in the presence of dynamic speckles induced by the turbulent atmosphere. Briefly, after low order AO correction to focus the majority of the photons in to the diffraction limited core and a coronagraph to reject the photons from the star there still remains some atmospheric turbulence induced speckles in the PSF halo in short exposures. In this speckle halo, at some locations the wavefront will interfere destructively and result in a zero photon event. As the atmosphere evolves and traverses the telescope field of view, the position of these nulls will change in the focal plane. However, at the position of a faint companion the probability of a zero photon event is considerably lower. Therefore, by counting the number of times each pixel records a zero photon event in each short exposure we can generate a `dark map' where the position of the companion will have a value lower than the rest of the image. Modern XAO systems are designed with target residual wavefront error (WFE) of the order of a few nm and so there is actually very little in the way of residual atmospheric speckle. However, as the quasi-static speckles are now dynamic we can use the dark speckle method to suppress the now dynamic quasi-static speckles. The DS method is intended to be used to enhance AO corrected images. Each DS exposure must be short otherwise the AO residual speckles will average, reducing the speckle nulls. Here we develop the technique for quasi-static speckles, in which case the exposure time no longer needs to be short and we record the minimum value over a number of mask configurations. An additional advantage of the pupil mask is that it is a configurable device allowing it to act in several different observing modes. In `adaptive' mode the mask can also be used to put a hard limit on the residual atmospheric WFE. Adaptive pupil masks like this have been shown to be able to reduce the PSF halo and actually increase PSF peak intensity despite removing photons \citep{Osborn09}, making it useful in scenarios where only low-order AO or even no AO is available. The APM could be used in `static' mode as a non-redundant aperture mask (e.g. \citeauthor{Kopilovich84}, 1984) or partially-redundant aperture mask (e.g. \citeauthor{Buscher93}, 1993), used in many modern high contrast imaging instruments. If the APM is sufficiently high order it can also be used to emulate any binary shaped pupil-plane mask coronagraph (e.g. \citeauthor{Nisenson01}, 2001; \citeauthor{Kasdin03}, 2003). The configuration of these static masks can easily be changed to experiment with different positions, element sizes and configurations, a current area of active research. In section 2 we describe the simulation, in section 3 we introduce the APM and show how it will reduce the quasi-static speckles, section 4 describes application of the dark speckle method to APM, in section 5 we discuss the results and we conclude in section 6.
|
Atmosphere induced phase aberrations (with or without AO) causes a halo of speckles to form around the PSF. This source of noise will average over time and, given enough time, faint companions can be observed above this halo. Current high contrast imaging is limited by quasi-static speckles caused by the optics and structure of the telescope. This is because these quasi-static speckles do not average. They appear in the image as potential false-positive candidates for faint companions and are difficult to distinguish. Here we have presented a technique to turn these static speckles into dynamic speckles. This means that over time these static speckles will also average into a broad halo. The shape and magnitude of this halo will depend upon the geometry of each mask element and the fraction of the pupil that is blocked. This means that over time the PSF will converge to a smooth form allowing the companions to be seen above. However, as the APM adds more energy to the halo and the throughput of the system is reduced by a fraction equal to the fraction of the pupil which is blocked, the {\it SNR} is reduced. We would need to observe for a longer period of time to collect enough photons from the companion to be seen above the halo. For an example case of a companion at $5\lambda/D$ we would need to observe for 1.35 times longer (1.47 defined by the simulation including noise) to achieve the same {\it SNR}. In this case the we do not need to expose the CCD between each mask state and so the mask frequency can be high allowing many mask states for an arbitrarily long exposure. In addition to the smoother PSF we can also use other conventional image manipulation techniques which are normally used for the dynamic speckling caused by atmospheric turbulence. Here we show the effect of implementing an adaptation of Labeyrie's dark speckle analysis. This is a simple technique where we record the minimum value of each pixel over all of the iterations of the mask. Using this technique in a Monte-Carlo simulation we can detect faint companions with a higher magnitude difference from the central star. We find that at separations greater than a few $\lambda/D$ the PSF halo count is reduced by a factor of approximately 100, corresponding to a contrast increase of 5 magnitudes. Due to the APM the diffraction rings from the bright star also move, this means that the diffraction pattern from the bright star is also suppressed by the APM when used with the DS imaging analysis. At small inner working angles there is still some confusion between the companions and that of the parent star, due to the pinned speckles in the bright diffraction rings. However, we are now able to detect companions 11 magnitudes fainter than the star at separations of $5\lambda/D$ and up to 18 magnitudes fainter at $22.5\lambda/D$. The inner working angle would be improved with the use of a coronagraph or other PSF subtraction methods which is not included here. In order to perform the DS analysis short exposure images will have to be recorded. The length of the short exposures is arbitrary but the optimal will depend on the number of iterations required during the exposure and the magnitude of the target. The noise attenuation found here is comparable with the attenuation from the High-Order Testbench (HOT) with ADI \citep{Martinez12}. The advantage of this technique over ADI is that it does not rely on PSF subtraction and is therefore insensitive to any changes in the quasi-static pattern. It has the advantage over PDI and SSDI that it also insensitive to the companion properties, i.e. it does not require the companion to have a specific emission spectrum or the light to be reflected and hence partially polarised. The disadvantage is the reduced throughput. APM would benefit from collaboration with other techniques. Currently the inner working angle is limited by the diffraction pattern of the primary star. A coronagraph would reject light from the central star resulting in even higher achievable contrast ratios. Also, no additional image manipulation was performed. No attempt at PSF subtraction was made. The dynamic speckles will average into the predictable smooth long exposure PSF, which could be subtracted. Combination with other techniques would improve the noise attenuation and reduce the probability of false-positives. It is also important to note that by including a high order APM, in addition to quasi-static speckle removal, it would also be possible to reduce the residual WFE, emulate a non-redundant (or partially redundant) aperture mask and a binary shaped pupil plane coronagraph, all of which would be completely and easily re-configurable.
| 12
| 6
|
1206.3222
|
1206
|
1206.1820_arXiv.txt
|
Recent work has shown that most globular clusters have at least two chemically distinct components, as well as cluster-to-cluster differences in the mean [O/Fe], [Mg/Fe], and [Si/Fe] ratios at similar [Fe/H] values. In order to investigate the implications of variations in the abundances of these and other metals for H-R diagrams and predicted ages, grids of evolutionary sequences have been computed for scaled solar and enhanced $\alpha$-element metal abundances, and for mixtures in which the assumed [$m$/Fe] value for each of the metals C, N, O, Ne, Na, Mg, Si, S, Ca, and Ti has been increased, in turn, by 0.4 dex {\it at constant} [Fe/H]. These tracks, together with isochrones for ages from $\approx 6$ to 14 Gyr, have been computed for $-3.0 \le$ [Fe/H] $\le -0.6$, with helium abundances $Y = 0.25$, 0.29, and 0.33 at each [Fe/H] value, using upgraded versions of the Victoria stellar structure program and the Regina interpolation code, respectively. Turnoff luminosity versus age relations from isochrones are found to depend almost entirely on the importance of the CNO-cycle, and thereby mainly on the abundance of oxygen. Since C, N, and O, as well as Ne and S, do not contribute significantly to the opacities at low temperatures and densities, variations in their abundances do not impact the predicted $\teff$ scale of red giants. The latter is a strong function of the abundances of only Mg and Si (and Fe, possibly to a lesser extent), because they are so abundant and because they are strong sources of opacity at low temperatures. For these reasons, Mg and Si also have important effects on the temperatures of MS stars. Due to their low abundances, Na, Ca, are Ti are of little consequence for stellar models. The effects of varying the adopted solar metals mix and the helium abundance at a fixed [Fe/H] are also briefly discussed.
|
\label{sec:intro} It is now a well-established result that the observed chemical abundance variations in globular cluster (GC) stars are due to both evolutionary processes within them and variations in the chemical makeup of the gas out of which they formed. There is no doubt that extra nonconvective mixing (not yet fully understood, but see \citealt{den12} for recent advances in our understanding) operates in cluster giants brighter than the red-giant-branch (RGB) bump (\citealt{dv03}, and references therein), which causes the surface abundances of C, N, and (sometimes) O to vary with luminosity. However, CN variations and the O--Na anticorrelation that appears to be characteristic of nearly all GCs (\citealt{cbg09}) have also been found in dwarf, turnoff, and subgiant stars (e.g., \citealt{ccb98}; \citealt{gbb01}; \citealt{rc02}; \citealt{cm05}; \citealt{dlg10}), where extra mixing is not a viable explanation. (NGC$\,$5466 appears to be a notable exception to this insofar as little or no evidence has been found in this system for primordial star-to-star variations in the abundances of the light elements; see \citealt{smw10}.) Indeed, below the RGB bump, the chemical composition anomalies do not appear to vary with luminosity (also see \citealt{gbn02}). Further evidence for primordial abundance variations in GCs is provided by the detection of Mg--Al anticorrelations (see \citealt{ygn05}). The stars currently on the main sequence (MS) and the RGB of these systems can hardly be responsible for the Al-rich, Mg-depleted (and, in a subset of the latter, Si-enhanced) stars that have been found in the most massive and/or the most metal-deficient clusters (like NGC$\,$2808, NGC$\,$6388, and M$\,$15 --- see \citealt{cbg09}) because their interiors are not hot enough for the Mg--Al cycle (and, especially, any additional processing to Si) to occur. As discussed by Carretta et al., the most likely explanation for the origin of the observed Mg, Al, and Si abundances is hot-bottom burning in intermediate-mass asymptotic-giant-branch (AGB) stars. In fact, there are cluster-to-cluster differences in the mean abundances of these elements as well as star-to-star differences within a given GC. For instance, as reported by Carretta et al., NGC$\,$2808, M$\,$4, and M$\,$5, which have similar [Fe/H] values to within $\approx 0.3$ dex, have mean [Mg/Fe] ratios of 0.20, 0.41, and 0.55, with {\it rms} variations of 0.25, 0.07, and 0.03 dex, respectively, while their [Si/Fe] values are, in turn, 0.28, 0.30, and 0.52, with {\it rms} variations amounting to $\approx 0.05$ dex in each cluster. Interestingly, the average [O/Fe] value in the Carretta et al.~sample of 17 GCs ranges from $-0.30$ in NGC$\,$6388 to $+0.46$ in NGC 7099, with {\it rms} variations from $\sim 0.1$ dex in several systems to as high as 0.36 dex (in NGC$\,$2808). These results raise at least two obvious questions. Is the net effect of the scatter in the [$m$/Fe] value for each metal $m$, within a given cluster, large enough to cause a detectable spread in the effective temperatures, and hence the colors, of stars along the principal sequences that define its color-magnitude diagram (CMD)? Are cluster-to-cluster differences in the abundances of individual metals big enough to affect the difference in color between the turnoff and the lower RGB, which is a sensitive measure of the relative ages of GCs having very similar [Fe/H] values (\citealt{vbs90}) and often used for that purpose (e.g., \citealt{map09})? \citet{vs91} showed early on that this relative age diagnostic is affected by differences in [O/Fe], but it is not yet clear whether it is also impacted by variations in the abundances of other $\alpha$-elements, such as Ne, Mg, and Si, which are the most abundant metals after the CNO group. Although there have been a few studies of metal-poor stellar models that allow for C--N--O--Na anticorrelations (notably \citealt{swf06}, \citealt{csp08}, \citealt{pcs09}), with some speculative discussion of the expected effects of Mg--Al anticorrelations (see the paper by Salaris et al.), the consequences of varying each of the most abundant metals, in turn, have yet to be adequately investigated. To be sure, we already have a very good understanding of the role played by C, N, and O in the evolution of lower mass stars and on computed isochrones (e.g., \citealt{rc85}, \citealt{van85}, \citealt{van92}\footnote{\citet{dcf07} incorrectly asserted that the models reported by VandenBerg (1992) ``did not account for enhanced oxygen in the opacities". In fact, opacity tables for temperatures $> 1$ eV ($\approx 12,000$ K) were derived for the assumed heavy-element abundances from the Los Alamos Astrophysical Opacity Library (\citealt{hmm77}), from which Rosseland mean opacities could be calculated for any mixture of the 20 most abundant elements. It was only the opacities for lower temperatures that were not obtainable at that time, but as mentioned therein and confirmed in the present investigation, oxygen does not contribute significantly to the low-$T$ opacities because it has a high ionization potential and thus is a poor electron donor at the temperatures and densities characteristic of the outer layers of stars (also see \citealt{bo86}, \citealt{vb01}).}). The same can be said for the case when the abundances of all of the $\alpha$-elements vary together (e.g., \citealt{scs93}, \citealt{vsr00}, \citealt{kdy02}, \citealt{pcs06}). However, relatively little has been done to date to examine the dependence of model $\teff$s on the abundances of individual $\alpha$-elements heavier than oxygen. The most noteworthy of the few available studies that have carried out such work is that by \citet[also see \citealt{lwd09}, who discuss some of the implications of the Dotter et al.~results for integrated colors, Lick indices, and synthetic specta]{dcf07}, but their analysis is complicated by the fact that their computations were carried out at a constant value of $Z$ (the mass-fraction abundance of all elements heavier than helium). This has the consequence that, when the abundance of one metal is increased, the abundances of all of the others are reduced. The net effect on stellar models therefore depends on how the surface and interior opacities have changed as a result of varying all of the elemental abundances at the same time, by different amounts. Fortunately, the problem is not as intractable as these few remarks suggest, because the [Fe/H] values that are obtained for the cases when the abundances of most of the metals (except oxygen) are enhanced, in turn, differ by $\lta 0.05$ dex from that of the base mixture. Hence, the effects of varying the chemical composition at a constant $Z$ is quite a good approximation to those derived when a single metal is enhanced at a constant value of [Fe/H]. Accordingly, the authors are able to confidently predict, for instance, that increased abundances of Mg and Si will result in cooler tracks and isochrones. Still, since observers use number abundance ratios to describe the chemical compositions of stars, stellar models that vary the abundances of a single element at a constant [Fe/H] value are much more straightforward to interpret. The latter approach, which is taken in the present study, has the advantage that the effects of altering the abundances of individual metals with can be accurately and precisely quantified. Moreover, whereas \citet{dcf07} considered only one (high) value of $Z$, we have generated large grids of tracks and isochrones for [Fe/H] values ranging from $-3.0$ to $-0.6$, in steps of 0.2 dex, on the assumption of $Y = 0.25$, 0.29, and 0.33 at each adopted iron abundance. At the lowest metallicities, one can anticipate that variations in the abundances of elements heavier than oxygen will have little or no impact on stellar evolutionary computations, as the interior opacity at low $Z$ is due primarily to bound-free and free-free processes involving H and He. The CNO elements, which are also considered here, differ from the heavier metals in that they affect the nucleosynthesis of hydrogen to helium, and thereby the structures of stars: an increase in the importance of the CNO-cycle due to enhanced abundances of these elements will make H-burning more centrally concentrated because of the high temperature sensitivity of these reactions compared to the $pp$-chain, and it is well known that the rate at which the radius of a low-mass star grows with time after leaving the zero-age main sequence (ZAMS) depends upon the rate at which its central concentration increases. These effects can be expected to be important at any metallicity.
|
\label{sec:sum} This investigation was carried out primarily to determine the impact on computed evolutionary tracks (for masses from 0.4 to $1.2 {{\cal M}_\odot}$) and isochrones (for ages in the range of 5--14 Gyr) of varying the abundances of each of several metals (specifically C, N, O, Ne, Na, Mg, Si, S, Ca, and Ti), in turn, at constant [Fe/H] values. Indeed, extensive model grids for each of these cases have been computed for [Fe/H] values ranging from $-3.0$ to $-0.6$, in 0.2 dex increments, with $Y = 0.25, 0.29$, and 0.33 at each [Fe/H] value. All of the tracks were generated using a significantly improved version of the Victoria stellar structure code, which now includes a treatment of the gravitational settling of helium and extra (turbulent) mixing very similar to the methods employed by \citet{pm91}, as well as the latest nuclear reaction rates and conductive opacities. Fully consistent OPAL and low-temperature Rosseland mean opacities were obtained, and used, for each of the adopted heavy-element mixtures. When very similar physics is assumed, the Victoria models are found to be in excellent agreement with those based on the MESA or the University of Montreal codes. Important improvements have also been made to the Regina interpolation code, which now produces isochrones, luminosity functions, etc., from a set of evolutionary tracks very efficiently using an interactive iterative interface. Moreover, it may be used to interpolate for arbitrary $Y$ and [Fe/H] values within the ranges for which grids of evolutionary tracks have been computed. We have verified that the interpolated grids for intermediate values of $Y$ and [Fe/H] reproduce those which are computed specifically for those abundance choices remarkably well, with respect to not only the placement of the tracks on the H-R diagram, but also the temporal derivatives of the luminosity and effective temperature along the tracks, which are important for the calculation of isochrone probability functions. Because they are poor electron donors at low temperatures (i.e., they do not contribute significantly to low-temperature opacities), C, N, O, Ne, and S do not affect the predicted location of the giant branch. However, as they are important sources of opacity at stellar interior conditions, tracks computed for enhanced abundances of these elements are somewhat cooler and fainter. C, N, and O (especially oxygen, because of its great abundance) have bigger effects on the turnoff-luminosity versus age relations from isochrones than any other metal mainly because of their role in the CNO cycle. Higher CNO abundances imply an increased importance of the CNO cycle and thereby more centrally concentrated burning which, in turn, has the ramification that the radius of a star grows more rapidly and the turnoff is reached sooner. By contrast, the effects on isochrones due to changes in the opacity are much less important. The remaining elements that were considered --- Na, Mg, Si, Ca, and Ti --- are important opacity sources at both low and high temperatures, but because of their low abundances, the effects of Na, Ca, and Ti are of no consequence (at least at low metallicities). As a result, only Mg and Si (and Fe, which is not given the same detailed analysis as the other elements) impacts the models for the RGB phase. Indeed, our study suggests that the location of the giant branch on the H-R diagram is a strong function of the abundances of Mg and Si, seemingly stronger than its dependence the iron content, which would not be too surprising given that the former are more abundant than the latter, especially in an $\alpha$-element enhanced metals mixture. Mg and Si also have significant consequences for the $\teff$ scale of MS and turnoff stars: only oxygen is more important for the turnoff phase of evolution. As opacity effects are largest at the coolest temperatures, and since O, Ne, and S affect the temperatures of MS and turnoff stars but not those of giants, variations in the abundances of O, Ne, Mg, Si, and S can all affect the difference in temperature (or color) between the turnoff and the lower RGB, when measured at some luminosity above the turnoff. As a result, it is risky to use this separation as a constraint on, in particular, the relative ages of star clusters having similar iron abundances (the method first proposed by VandenBerg et al.~1990), unless the abundances of these 5 metals have been determined and the effects of cluster-to-cluster differences in them taken into account. Star-to-star variations in the abundances of Mg and Si may also be responsible for the broadening of the RGB in some GCs, and spreads in the abundances of these metals, plus O, Ne, and S, could potentially affect the width of the MS. As long as it is clear which solar abundances (e.g., GS98, AGS5, A09) have been used in deriving the [$m$/Fe] values from the observed $m$/H number-abundance ratios in metal-poor stars, and the correct [$m$/Fe] values are assumed in the computation of the isochrones which are fitted to globular cluster CMDs, the ages so obtained should be independent of this choice. Importantly, the turnoff luminosity versus age relations from isochrones are primarily a function of the oxygen and helium abundances. Finally, we note that it is our intention to carry out a supplementary study to address a number of unanswered questions. Besides carrying out some comparisons of the observed widths of the principal photometric sequences of globular clusters with models that allow for varying abundances of several metals (notably O, Mg, and Si), the forthcoming work will isolate and examine the effects of varying the Fe abundance (keeping all other abundances fixed), and it will determine, for cases of particular interest (including Ne--Na and Mg--Al anticorrelations), the impact of varying the abundances of two elements at the same time. Such work will shed some light on whether the net effect of varying the abundances of several elements is equivalent to adding the effects due to each individual element, and it will tell us, for instance, whether the tracks and isochrones that are computed for normal Mg and Al abundances reproduce those for reduced Mg and the significantly enhanced Al abundance that would be expected (because the latter is initially much less abundant than the former) if the sum of the Mg and Al abundances is constant. (It is our intention to make the model grids that will be computed in the forthcoming study, along with a subset of those generated for the present investigation, generally available through the Canadian Astronomical Data Center. Those with a more immediate need for some of the computations reported herein are asked to contact the first author.)
| 12
| 6
|
1206.1820
|
1206
|
1206.4919_arXiv.txt
|
The identification of strong gravitational lenses in large surveys has historically been a rather time consuming exercise. Early data from the \emph{Herschel} Astrophysical Terahertz Large Area Survey (\emph{Herschel}-ATLAS) demonstrate that lenses can be identified efficiently at submillimetre wavelengths using a simple flux criteria. Motivated by that development, this work considers the statistical properties of strong gravitational lens systems which have been, and will be, found by the \emph{Herschel}-ATLAS. Analytical models of lens statistics are tested with the current best estimates for the various model ingredients. These include the cosmological parameters, the mass function and the lens density profile, for which we consider the singular isothermal sphere (SIS) and the Navarro, Frenk \& White (NFW) approximations. The five lenses identified in the \emph{Herschel}-ATLAS Science Demonstration Phase suggest a SIS density profile is preferred, but cannot yet constrain $\oml$ to an accuracy comparable with other methods. The complete \emph{Herschel}-ATLAS data set should be sufficient to yield competitive constraints on $\oml$. Whilst this huge number of lenses has great potential for constraining cosmological parameters, they will be most powerful in constraining uncertainty in astrophysical processes. Further investigation is needed to fully exploit this unprecedented data set.
|
Since the first strong gravitational lens was discovered in 1979 (\citeauthor{Walsh1979}), hundreds have been found and studied \citep{Treu2010}. As well as serendipitous discoveries, dedicated surveys have searched for lens candidates. The Cosmic Lens All-Sky Survey \citep{Browne2003} identified 22 new strong gravitational lenses from imaging in the radio waveband. The Sloan Lens ACS Survey \citep{Bolton2006,Auger2009} has so far confirmed 85 new gravitational lenses using spectroscopic data from the SDSS to identify candidates for follow-up with high-resolution HST imaging. The Cosmological Evolution Survey \citep{Faure2008,Jackson2008} identified 67 lenses from visual inspection of HST imaging. In addition there are projects, such as MUSCLES \citep{Jackson2011} and SQLS \citep{Inada2012}, that are looking for new lenses by reprocessing data from large surveys such as UKIDSS and SDSS. There has been much interest in studying gravitational lenses because they magnify sources that would otherwise be too faint or distant to see. Historically these lens systems have been difficult to identify in sufficiently large numbers to be statistically useful. Some lens searches involve painstaking `by eye' analysis of each object observed. Now new methodologies are being developed to improve the efficiency of finding lenses and to improve the conversion rate from candidate to confirmed lensed source. These include using colours \citep{Ofek2007}, spectroscopic redshifts \citep{Auger2009}, and other imaging processing techniques \citep{Marshall2008}. For a rigorous statistical study of gravitational lenses, they should all come from a single survey. This is because the probability of identifying lensed sources varies between surveys due to the different instrumental sensitivities and selection effects. For sources observed in the submillimetre waveband it has long been thought that galaxy number counts should fall off rapidly with increasing flux density. This is because of the rapid evolution of proto-spheroidal galaxies into starburst galaxies \citep{Granato2004}. This drop in counts with flux using submillimetre observations was anticipated to yield an effective way of identifying lenses \citep{Blain1996,Negrello2007}. It has since been successfully demonstrated using the \emph{Herschel}-ATLAS data \citep{Negrello2010, Gonzalez-Nuevo2012}. Lensed sources are identifiable much more easily because the lensing magnification pushes them into higher flux densities which have intrinsically low source populations. The \emph{Herschel} Space Observatory, launched in 2009, is the only space-based observatory to cover a spectral range from the far infrared to the submillimetre. It provides a unique window through which to study large scale structure. The \emph{Herschel} Astrophysical Terahertz Large Area Survey \citep[\emph{Herschel}-ATLAS]{Eales2010} will cover 550 deg$^2$ of the sky making it the \emph{Herschel} survey with the largest area. \cite{Negrello2010} showed that strong gravitational lenses could be identified with almost 100\% efficiency by selecting sources which have 500$\mu$m flux densities greater than 100mJy. \cite{Gonzalez-Nuevo2012} further improved on this technique by refining the flux selection to incorporate other \emph{Herschel} wavebands and optical/infrared data from the VIKING \citep{Fleuren2012,Sutherland2012} survey to look for counterparts. This technique is forecast to yield ten times more lenses than that by \citeauthor{Negrello2010}. The expected efficiency of 70\% is lower that that achieved by \citeauthor{Negrello2010}, but still impressive. Using these techniques, the \emph{Herschel}-ATLAS team expect to find as many as a thousand strongly lensed candidates. We emphasise that this article considers the likelihood of \emph{strong} gravitational lenses which are defined as lensed systems with multiply imaged sources. Recent articles on the effect of lensing on the source number counts \citep{Lima2010, Wardlow2012} do not consider strong lensing in the sense that sources are multiply imaged, just that they are highly magnified. Note, this results in the lens probability distribution assuming a NFW density profile becoming much more similar to that of the SIS density profile than is seen in our analysis. Looking to the future, `strongly' lensed sources are being selected by their excess submillimetre flux, or effectively by their magnification, with limited optical follow-up to confirm multiple imaging. Since requiring optical follow-up may limit the number of confirmed sources, it may in future turn out to be statistically favourable to relax the multiple image requirement to gain a statistical advantage. However our approach, looking at strong gravitational lens systems with multiple images, is used to be consistent with the nature of the sources identified by \cite{Negrello2010}. The statistical properties of gravitationally lensed sources were first considered in the 1980s \citep{Turner1984} and later applied as a possible way of constraining the cosmological model \citep{Li2002,Mitchell2005,Zhang2009,Dobke2009}. With the prospect of a thousand lenses being identified by \emph{Herschel}-ATLAS, we take a fresh look at the analytical theory behind predicting strong gravitational lens statistics in Sections \ref{sec1a} and \ref{sec1b}. In Section \ref{sec2} the lenses identified by \cite{Negrello2010} are compared against the analytical predictions, and then in Section \ref{sec3} we consider the parameter constraints possible with the full \emph{Herschel}-ATLAS data set. Unless otherwise stated we use cosmological parameters: $\omm$ = 0.27, $\oml$ = 0.73, $\Omega_\textsubscript{b}$ = 0.046, h = 0.7, n = 0.97, and $\sigma_8$ = 0.81 \citep{Komatsu2011}.
|
Until recently the number of strong gravitational lenses that had been identified was relatively small for any particular survey. Now this situation is about to change dramatically as future surveys are expected to yield numbers of strong lenses into the hundreds and thousands. One of these surveys, \emph{Herschel}-ATLAS, expects to yield up to 1000 lenses. In this article we considered the statistics that can be done with this new data set of strong gravitational lenses. First we reviewed the current analytical theory. We noted that the standard method for calculating magnification bias was not appropriate here. This is because the fast fall in number counts with flux in the submillimetre wavebands means that the standard calculation gives an infinitely large magnification bias. We therefore proposed a more realistic method. This effectively means the magnification bias is equal to the mean expected magnification of the gravitational lenses. We saw that lensing statistics are sensitive to astrophysical properties, as well as the cosmological parameters, most significantly the lens density profile. Although so far only five lenses have been confirmed in the \emph{Herschel}-ATLAS SDP data, limiting its the statistical value, we have shown that these can be used to constrain the lens density profile. Initially the likelihoods calculated from each of $\pmu$ and $\pzl$ suggested conflicting results between the SIS and NFW lens density profiles. But simulations of the expected uncertainty of these likelihoods due to the small number of lenses showed that the results are perfectly consistent with a SIS, not a NFW, density profile. Of course neither of the SIS or NFW density profiles is probably realistic to explain the complex structure of the dark matter halo and baryon component. Here they are used to illustrate the effect of the current uncertainty in the density profile on the resulting optical depth. We considered just one of the cosmological parameters for illustrative purposes. Whilst five lenses is not yet enough information to constrain the dark energy density, the full \emph{Herschel}-ATLAS data set should be more than sufficient to provide constraints competitive with other methods. How the mass function could be independently constrained was more uncertain and further work is clearly required to understand how this and other other uncertainties can most effectively be constrained. Here we have seen that the dominant uncertainties in gravitational lens statistics are not cosmological but astrophysical i.e. understanding the nature of the lensing objects. Whilst more investigation is required, we can see that new data sets of strong gravitational lenses will be fundamental for constraining both halo astrophysics as well as cosmological parameters.
| 12
| 6
|
1206.4919
|
1206
|
1206.0826_arXiv.txt
|
{We present a 1.1mm emission map of the OMC1 region observed with AzTEC, a new large-format array composed of 144 silicon-nitride micromesh bolometers, that was in use at the James Clerk Maxwell Telescope (JCMT). These AzTEC observations reveal dozens of cloud cores and a tail of filaments in a manner that is almost identical to the submillimeter continuum emission of the entire OMC1 region at 450 and 850 $\mu$m. We perform Fourier analysis of the image with a modified periodogram and the density power spectrum, which provides the distribution of the length scale of the structures, is determined. The expected value of the periodogram converges to the resulting power spectrum in the mean squared sense. The present analysis reveals that the power spectrum steepens at relatively smaller scales. At larger scales, the spectrum flattens and the power law becomes shallower. The power spectra of the 1.1mm emission show clear deviations from a single power law. We find that at least three components of power law might be fitted to the calculated power spectrum of the 1.1mm emission. The slope of the best fit power law, $\gamma \approx -2.7$ is similar to those values found in numerical simulations. The effects of beam size and the noise spectrum on the shape and slope of the power spectrum are also included in the present analysis. The slope of the power law changes significantly at higher spatial frequency as the beam size increases.}
|
The Orion Nebula is located in the northern part of the Orion A molecular cloud and consists of OMC1, OMC2, OMC3, and OMC1-S. These components are smoothly connected and form integral shaped filaments seen in continuum observations at sub-millimeter, infrared, and optical wavelengths. The M42, Orion Nebula, is known to be the best studied star-forming region in the nearest giant molecular cloud (GMC) where massive stars have formed and has been extensively studied at all wavelengths. OMC1 in the Orion A molecular cloud possesses the most massive component containing at least 2,000 young stellar objects (YSOs) (O'Dell 2001) and is strongly connected to the surrounding gas. There is also a report on the VLA NH$_3$ observations and revelations of the filamentary and clumpy structure of OMC1 in the Orion A molecular cloud (Wiseman \& Ho 1996, 1998). OMC1 has an on-going massive star formation and strong interactions between gas and young stellar objects (YSOs), their bipolar stellar jets, and outflows (Mann \& Williams 2010; Bally et al. 2011). As we have mentioned above, the complex and filamentary structure of the Orion A molecular cloud, especially the Orion Integral Shaped Filament was revealed by sub-millimeter continuum observations by Lis et al. (1998) and Johnstone \& Bally (1999). The condensation mass spectrum was also explored and temperature distribution was examined (Johnstone \& Bally 1999). Mookerjea et al. (2000) found that the coldest clump in the Orion A molecular cloud had a temperature of about 15 K. These cold clumps can be traced at the continuum observations using the observations at mm wavelength. Despite various observations of the Orion A molecular cloud over a wide range of wavelengths for over a decade including the SEST SIMBA (Nyman et al. 2001), thermal continuum emission at 1.1mm has been observed for the first time only recently, owing to the development of the Astronomical Thermal Emission Camera (AzTEC). This AzTEC camera is composed of an array of 144 nitride micro-mesh composite bolometers (Wilson et al. 2008). This camera was originally designed for a millimeter-wavelength bolometer camera to be installed on the Large Millimeter Telescope (LMT). But in year 2005 and 2006, it was installed on the James Clerk Maxwell Telescope (JCMT) in Hawaii, Mauna Kea and performed biased and unbiased surveys, of the northern sky. The present study analyzes the 1.1mm AzTEC observations of the OMC1 components and reports the fact that the 1.1mm emission from OMC1 is distributed in a manner almost similar to that of the 850 $\mu$m emission (Johnstone \& Bally 1999) observed with the Submillimeter Common-User Bolometer Array (SCUBA) on the JCMT. In the present analysis, we performed a power spectrum analysis of the 1.1mm intensity map which is dominated by thermal emission from cold dust in OMC1 in the Orion A molecular filaments. Since the dust grains are coupled with gas, elucidating the structure of the dust emission will also play a crucial role in understanding the dynamics and the structure of the gas. In general, the density fluctuations of the interstellar dust (Draine \& Lazarian 1998) occur in the cold and dense interstellar medium (ISM). The power spectrum analysis using the Fourier conversion of the given images is especially important because they can reveal the statistical properties of structures which may be present across the images. The generation of the hierarchical structures or fractal structures (Falgarone et al. 1991; Elmegreen \& Falgarone 1996) often presents across a range of spatial scales in the ISM. It has also been suggested that this could be attributed to turbulent emission since this can increase gas density and initiate star forming instabilities (Mac Low 2004; Burkert 2006). However, the large scale properties of turbulence in the ISM, including those of the gas and dust grains are still poorly understood. But there has been an explosion of research in this area. In general, the energy input from star formation is thought to be a major driving force for the turbulence in the ISM (Scalo \& Elmegreen 2004). It is also well known that measurements of the slopes of the power spectra (Crovisier \& Dickey 1983; Gautier et al. 1992) from the intensity fluctuations in the turbulent medium can provide an insight into the nature of scales present in the hierarchical or non-hierarchical structures (Blitz \& Williams 1997; Hartmann 2002) in the interstellar medium (Brunt 2010). Thus, the results from the present studies of the power spectrum analysis can reveal the nature of the structures seen in the 1.1mm image of OMC1 in the Orion A molecular cloud and filaments. The results of this present analysis are presented in Section 3 and 4. Section 2 provides procedures of image processing and the resultant data.
|
We present a 1.1mm emission map of the OMC1 region observed with AzTEC, a new large-format array composed of 144 silicon-nitride micromesh bolometers that was in use at the JCMT. We performed a Fourier analysis of the image with a modified periodogram. The periodogram was measured by computing the discrete Fourier transformation. From the present analysis of the OMC1 filaments at 1.1mm emission, the power spectrum steepens at relatively smaller scales. At larger scales, the power spectrum flattens and the large scale power law becomes shallower. The logarithmic slopes of the power-law spectra obtained seem to be lower than the logarithmic slope of Kolmogorov spectrum. We also performed power spectrum calculations using three different aperture diameters. As the size of aperture increases, the power decreases significantly. We were able to fit at least three components of power law in the power spectrum of the 1.1mm emission map. The slope of the best fit at spatial frequency of $30 \lesssim k \lesssim 100$ is $\gamma\approx-2.66\pm0.3$. This result is similar to the spectral index of the power spectrum, $\gamma\approx-2.7$ that was found in numerical simulations. The effects of noise on the slope of the power spectrum were also included in the present analysis.
| 12
| 6
|
1206.0826
|
1206
|
1206.5542_arXiv.txt
|
Neutron stars enable us to study both the highest densities and the highest magnetic fields in the known Universe. In this article I review what can be learned about such fundamental physics using magnetar bursts. Both the instability mechanisms that trigger the bursts, and the subsequent dynamical and radiative response of the star, can be used to explore stellar and magnetospheric structure and composition.
|
Neutron stars have densities so high that nuclei, and even neutrons, may dissolve to form exotic states of matter. The high densities also allow neutron stars to sustain magnetic fields ten orders of magnitude higher than those we can create in terrestrial facilities, in regimes where unusual electromagnetic processes are expected. These properties allow us to explore physics that cannot be studied in the laboratory. \subsection{Dense matter} The theory of Quantum Chromodynamics provides an excellent description of how quarks are bound together to form nucleons. Understanding the interaction between nucleons, however, is much more complex. Many-body nucleon theory (even at the level of the three-nucleon interaction), particularly for isospin asymmetric matter and for densities exceeding the saturation density, is especially challenging. The nature of this interaction, and the state of matter at extremes of temperature and density, are extremely active fields of research whose resolution requires an integrated effort between collider experiments and relativistic astrophysics. Of particular interest is the possibility of transitions from nucleons to de-confined quarks and gluons or other more exotic states. At low densities this can be studied by experiments like the Large Hadron Collider or in various Heavy Ion experiments such as the Relativistic Heavy Ion Collider and the Facility for Antiproton and Ion Research. At higher densities, however, neutron stars are the only environment in the Universe where the transition can be explored. \subsection{Strong magnetic fields} The strongest magnetic fields that can be studied within our Solar System are those created in high magnetic field laboratories on Earth. At present these facilities can generate fields approaching $10^6$ G, but only for fractions of a second. Neutron stars are the only objects that let us study fields above $10^9$ G, and are the only stars with fields exceeding the quantum critical limit of $B_\mathrm{QED} = 4.4\times10^{13}$ G. At this point the magnetic field energy is so high that it can generate electron-positron pairs, so that the vacuum seethes with charged particles. In such an environment we expect a host of unusual and intriguing physical effects such as spontaneous photon splitting and vacuum birefringence.
|
Violent dynamical events that shake up the neutron star can be useful tools to study both the composition of the star (which depends on highly uncertain nuclear physics) and strong magnetic fields. For magnetar bursts, however, there are many open theoretical questions motivated by a wealth of excellent observational data. The way in which stress builds up in the system and the nature of the trigger for the bursts remains unknown. The dynamical response of the star, which appears to include the possibility of exciting global seismic vibrations, is far more complex than originally envisaged. However the possibility of identifying robust signatures of magnetic fields above the quantum critical limit, and of being able to use seismology to study neutron star interiors, are strong motivating factors for the determined astrophysicist.
| 12
| 6
|
1206.5542
|
1206
|
1206.0547_arXiv.txt
|
{ We present a photometric study of the globular clusters (GCs) in the Virgo giant elliptical galaxy M86 based on Washington $CT_1$ images. The colors of the GCs in M86 show a bimodal distribution with a blue peak at $(C-T_1)=1.30$ % and a red peak at $(C-T_1)=1.72$. % The spatial distribution of the red GCs is elongated similarly to that of the stellar halo, while that of the blue GCs is roughly circular. The radial number density profile of the blue GCs is more extended than that of the red GCs. The radial number density profile of the red GCs is consistent with the surface brightness profile of the M86 stellar halo. The GC system has a negative radial color gradient, which is mainly due to the number ratio of the blue GCs to the red GCs increasing as galactocentric radius increase. The bright blue GCs in the outer region of M86 show a blue tilt: the brighter they are, the redder their mean colors get. These results are discussed in comparison with other Virgo giant elliptical galaxies. }
|
Thousands of globular clusters (GCs) are found in a giant elliptical galaxy (gE) and they are found to be located from the center to the outer halo of their host galaxy \citep{lee03,bro06}. Therefore, the GCs in gEs are a powerful tool to study the structure and evolution of the GC system itself as well as their host galaxy. M86 (NGC 4406, VCC 881) is a famous gE located close to the center region of the Virgo galaxy cluster in the sky and is known to be infalling to the Virgo center with the relative velocity of about 1300 % km s$^{-1}$. % It is also known early to be abundant with GCs \citep{han77}. There are several photometric studies for the GCs in M86 in the literature. \citet{coh88} presented $gri$ photometry of GCs at $R<7\arcmin$ in M86 obtained at the Hale 5 m telescope. She found that the GC system of M86 is similarly extended as the stellar light, and that there is no detectable radial color gradient. % From HST/WFPC2 $VI$ observation of the central region of M86, \citet{nei99} reported that the color distribution of the GCs in M86 shows a single peak. However, \citet{kun01} and \citet{lar01} showed later, using the same data as used in HST/WFPC2 by \citet{nei99}, that the color distribution of the GCs in M86 is bimodal. Later \citet{pen06} confirmed, from HST/ACS $gz$ observation, that the color distribution of M86 GCs is bimodal. \citet{jor07} and \citet{vil10} derived a luminosity function of M86 GCs from the same HST/ACS images, and obtained a similar Gaussian peak value at $g$-band, $\mu_g=23.950\pm0.097$ and $23.887\pm0.087$, respectively. \citet{pen08} estimated the specific frequency of M86 GCs, $S_N=2.57\pm0.12$, from the HST/ACS images. On the other hand, \citet{rho04} investigated M86 GCs using the wide field ($36\arcmin\times36\arcmin$) $BVR$ images obtained using the KPNO 4 m Mayall telescope. They found again that M86 GCs show a bimodal color distribution and that they show a modest negative color gradient with galactocentric radius. They derived a specific frequency of $S_N=3.5\pm0.5$, which is larger than the value derived from the central region by \citet{pen08}. However, previous studies based on HST data covered only a small central region of M86 and those based on ground-based data did not investigate the properties of sub-populations in the GC system and the spatial distribution of the GCs in M86. Here we investigate the detailed properties of the M86 GC system % using Washington photometry derived from deep and wide CCD images. The Washington filter system with a wide bandwidth is known to be very sensitive to measuring the metallicity of the GCs \citep{gei90} so that it is useful for studying the properties of sub-populations of the GCs. We adopt a distance to M86, derived from the surface brightness fluctuation method by \citet{mei07}, 16.9 Mpc ($(m-M)_0 =31.13\pm0.07$). At this distance one arcsec corresponds to a linear scale of 81 pc. The basic information of M86 is listed in Table \ref{tab-info}. This paper is composed as follows. Section 2 describes observation and data reduction. In \S 3, we present the color-magnitude diagram and color distribution of the M86 GCs. We investigate the spatial distribution of the GC system as well as the radial variation of number densities and colors of the GCs. In \S 4, we discuss our results and their implication in comparison with other Virgo gE studies. Primary results are summarized in \S 5. \begin{deluxetable}{lcc} \tablecaption{Basic information of M86\label{tab-info}} \tablewidth{0pt} \tablehead{ \colhead{Parameters} & \colhead{Values } & \colhead{References} } \startdata R.A., Decl. (J2000) & $12^h~26^m~11.743^s$, +12$\arcdeg$ 56$\arcmin$ 46.40$\arcsec$ & 1 \\ Total magnitudes & $V^T=8.90\pm0.05$, $B^T=9.83\pm0.05$ & 2\\ Foreground reddening, $E(B-V)$ & 0.030 & 3\\ Distance, $d$ & 16.86 Mpc ($(m-M)_0=31.13\pm0.07$) & 4 \\ Systemic radial velocity, $v_p$ & $-244\pm5$ km s$^{-1}$ & 5\\ Effective radius, $R_{eff}$ & 3.59 arcmin ($C$), 3.14 arcmin ($T_1$) & 6\\ Effective ellipticity, $\epsilon_{eff}$ & 0.36 ($C$), 0.33 ($T_1$) & 6 \\ Effective P.A., $\Theta_{eff}$ & 119 deg ($C, T_1$) & 6 \\ Standard radius, $R_{25}$ & 4.88 arcmin ($C$), 9.64 arcmin ($T_1$) & 6\\ Standard ellipticity, $\epsilon_{25}$ & 0.40 ($C, T_1$) & 6 \\ Standard P.A., $\Theta_{25}$ & 124 deg ($C, T_1$) & 6 \\ \enddata \tablerefs{(1) NASA Extragalactic Database; (2) \citet{dev91}; (3) \citet{sch98}; (4) \citet{mei07}; (5) \citet{smi00}; (6) This study. } \end{deluxetable}
|
\begin{figure}[!t] \plotone{fig12.eps} \caption{ Comparison of the color distribution between the GCs in M86 and the GCs in other Virgo gEs. (a) M86 from this study. (b) M60 from \citet{lee08}. (c) M49 from \citet{gei96} and \citet{lee98}. The solid and dashed lines represent the double Gaussian fits. \label{fig-gE3col}} \end{figure} We compared the photometric results of the M86 GC system with those of GC systems in other two gEs in Virgo (M60 and M49). M60, which is the third brightest gE in Virgo and located about 3 deg from the Virgo center, and M49, which is the brightest gE in Virgo and located about 4 deg from the Virgo center. Previous studies for the GC systems of M60 \citep{lee08} and M49 \citep{gei96b,lee98} used the same instruments and similar photometric techniques as this study for the M86 GC system. For comparison we used the same criteria for the GC sample in three galaxies, GCs with $R<7\arcmin$ and $19<T_1<23$. Figure \ref{fig-gE3col} displays the color distributions of the GCs in M86, M60, and M49. The GCs in these galaxies show similarly bimodal color distributions, but there are slight differences among these galaxies. The peak colors for M86 ($(C-T_1)=1.31$ and 1.74) are similar to those for M49 ($(C-T_1)=1.30$ and 1.79), but are about 0.1 mag bluer than those for M60 ($(C-T_1)=1.37$ and 1.87). Applying this color difference to the relation between $(C-T_1)$ color and metallicity \citep{lee08}, this implies that the M86 and the M49 GCs are on average about 0.2 dex more metal-poor than M60 GCs. These results are approximately consistent with those by \citet{pen06} who covered the central regions with $R\lesssim 1\arcmin.5$. They reported that the peak colors for these three galaxies are $(g-z)=0.98$ and 1.33 for M86, $(g-z)=0.97$ and 1.42 for M49, and $(g-z)=0.98$ and 1.45 for M60. There are significant differences among the number ratios (N(BGC)/N(RGC)) of the blue GCs to the red GCs in these galaxies: $2.37\pm 0.26$ for M86, $1.38\pm 0.11$ for M60, and $0.97\pm 0.07$ for M49. Thus the fraction of the blue GCs for M86, the most faint among these three galaxies, is the largest. This is consistent with previous finding that the fraction of the blue GCs decreases as their host galaxy gets brighter \citep{pen06}. On the other hand, the number ratio of the blue GCs to the red GCs in each galaxy is a little bit larger than the value by \citet{pen06} ( $2.33\pm 1.15$ for M86, $0.75\pm 0.10$ for M60, and $0.69\pm 0.09$ for M49) and by \citet{fai11} ($1.07\pm 0.10$ for M60). The radial coverages used by \citet{pen06} and \citet{fai11} are $R\lesssim 1\arcmin.5$ and $R\lesssim 5\arcmin$, respectively, so that their fields are on average closer to the galaxy center than that used in this study ($1\arcmin<R<7\arcmin$). These shows that % this number ratio decreases as the galactocentric radius decreases (see Section 3.2 for detail). Recent studies for the GCs in gEs based on HST/ACS data \citep{har06,str06, mie06, pen09, mie10} and wide-field images \citep{for07, lee08, har09}, found a so-called `blue tilt' (or a color-magnitude relation): the brighter the bright blue GCs are, the redder their colors get. Several mechanisms were proposed to explain the origin of the blue tilt \citep{mie06, bek07, str08,har09b,bla10}. The self-enrichment process among these appears to be a main driver \citep{mie06, str08}. We investigated any existence of this blue tilt in the M86 GCs. Figure \ref{fig-gE3cmd} displays the CMDs of M86, M60, and M49 derived from the $CT_1$ images for the outer regions at $1\arcmin.5 <R <7\arcmin$. The blue and red peak values are determined from the KMM test with the colors in each magnitude bin. In Figure \ref{fig-gE3cmd} the blue tilt is seen for the bright blue GCs in M86 as well as other two galaxies. However, the number of galaxies is only three so that it needs to study more galaxies to address the slope difference of the blue tilt. The existence of the blue tilt in M86 as well as M49 and M60 is consistent with the result for high mass galaxies in the ACSVCS given by \citet{mie10}. The spatial distribution of the GCs in three galaxies share common features as follows: (a) the radial number density profile of the red GCs in the outer region is consistent with the surface brightness profile of the stellar halo, (b) the elongation of the red GC system is consistent with that of the stellar halo, while the spatial distribution of the blue GC system is approximately circular, (c) the mean color of the red GCs is similar to that of the stellar halo. These common properties are also shown for other gE GC systems such as M87 and NGC 1399 \citep{bas06,tam06b,for07,fai11,str11,for12}. Thus % the red GCs are more correlated with the stellar halo than the blue GCs. This indicates that most red (metal-rich) GCs in gEs might have formed with the halo stars.
| 12
| 6
|
1206.0547
|
1206
|
1206.5006_arXiv.txt
|
Various formulations of smooth-particle hydrodynamics (SPH) have been proposed, intended to resolve certain difficulties in the treatment of fluid mixing instabilities. Most have involved changes to the algorithm which either introduce artificial correction terms or violate what is arguably the greatest advantage of SPH over other methods: manifest conservation of energy, entropy, momentum, and angular momentum. Here, we show how a class of alternative SPH equations of motion (EOM) can be derived self-consistently from a discrete particle Lagrangian -- guaranteeing manifest conservation -- in a manner which tremendously improves treatment of these instabilities and contact discontinuities. Saitoh \&\ Makino recently noted that the volume element used to discretize the EOM does not need to explicitly invoke the mass density (as in the ``standard'' approach); we show how this insight, and the resulting degree of freedom, can be incorporated into the rigorous Lagrangian formulation that retains ideal conservation properties and includes the ``$\nabla h$'' terms that account for variable smoothing lengths. We derive a general EOM for any choice of volume element (particle ``weights'') and method of determining smoothing lengths. We then specify this to a ``pressure-entropy formulation'' which resolves problems in the traditional treatment of fluid interfaces. Implementing this in a new version of the {\small GADGET} code, we show it leads to good performance in mixing experiments (e.g. Kelvin-Helmholtz \&\ ``blob'' tests). And conservation is maintained even in strong shock/blastwave tests, where formulations without manifest conservation produce large errors. This also improves the treatment of sub-sonic turbulence, and lessens the need for large kernel particle numbers. The code changes are trivial and entail no additional numerical expense. This provides a general framework for self-consistent derivation of different ``flavors'' of SPH.
|
\label{sec:intro} Smoothed particle hydrodynamics (SPH) is a method for solving the equations of hydrodynamics \citep[in which Lagrangian discretized mass elements are followed;][]{lucy:1977.sph,gingold.monaghan:1977.sph} which has found widespread application in astrophysical simulations and a range of other fields as well \citep[for recent reviews, see][]{rosswog:2009.sph.review,springel:2010.sph.review,price:2012.sph.review}. % The popularity of SPH owes to a number of properties: compared to many other methods, it is numerically very robust (stable), trivially allows the tracing of individual fluid elements (Lagrangian), automatically produces improved resolution in high-density regions without the need for any ad-hoc pre-specified ``refinement'' criteria (inherently adaptive), is Galilean-invariant, couples properly and conservatively to N-body gravity schemes, exactly solves the particle continuity equation,\footnote{This is the continuity equation for a {\em discretized} particle field. Exactly solving the continuity equation for a continuous fluid, of course, requires infinite resolution or infinite ability to distort the Lagrangian particle ``shape.''} and has excellent conservation properties. The latter character stems from the fact that -- unlike Eulerian grid methods -- the SPH equations of motion (EOM) can be rigorously and exactly derived from a discretized particle Lagrangian, in a manner that guarantees manifest and simultaneous conservation of energy, entropy, linear momentum, and angular momentum \citep[][henceforth \Sentropy]{springel:entropy}. However, there has been considerable discussion in the literature regarding the accuracy with which the most common SPH algorithms capture certain fluid mixing processes \cite[particularly the Kelvin-Helmholtz instability; see e.g.][]{morris:1996.sph.stability,dilts:1999.sph.stability,ritchie.thomas:2001.egy.wtd.sph,marri:2003.mod.sph.cosmo.sims,okamoto:2003.shear.sph.flows,agertz:2007.sph.grid.mixing}. Comparison between SPH and Eulerian (grid) methods shows that while agreement is quite good for super-sonic flows, strong shock problems, and regimes with external forcing (e.g.\ gravity); ``standard'' SPH appears to suppress mixing in sub-sonic, thermal pressure-dominated regimes associated with contact discontinuities \citep{kitsionas:2009.grid.sph.compare.turbulence,price:2010.grid.sph.compare.turbulence,bauer:2011.sph.vs.arepo.shocks,sijacki:2011.gadget.arepo.hydro.tests}.\footnote{In fairness, we should emphasize that it has long been well-known that Eulerian grid codes, on the other hand, err on the side of {\em over}-mixing (especially when resolution is limited), and in fact this problem actually motivated some of the SPH work discussed above. This may, however, be remedied in moving-mesh approaches (though further study is needed; see e.g.\ \citealt{springel:arepo}).} The reason is, in part, that in standard SPH the kernel-smoothed density enters the EOM, and so behaves incorrectly near contact discontinuities (introducing an artificial ``surface tension''-like term) where the density is not differentiable. A variety of ``flavors'' (alternative formulations of the EOM or kernel estimators) of SPH have been proposed which remedy this \citep[see above and][]{monaghan:1997.sph.drag.viscosities,ritchie.thomas:2001.egy.wtd.sph,price:2008.sph.contact.discontinuities,wadsley:2008.sph.mixing.cosmology,read:2010.sph.mixing.optimization,read:2012.sph.w.dissipation.switches,abel:2011.sph.pressure.gradient.est,garciasenz:2012.integral.sph}. These approaches share an essential common principle, namely recognizing that the pressure at contact discontinuities must be single-valued (effectively removing the surface tension term). Some of these show great promise. However, many (though not all) of these formulations either introduce additional (potentially unphysical) dissipation terms and/or explicitly violate the manifest conservation and continuity solutions described above -- perhaps the greatest advantages of SPH. This can lead to severe errors in problems with strong shocks or high-Mach number flows, limited resolution, or much larger gradients between phase boundaries (J.\ Read, private communication; see also the discussion in \citealt{price:2012.sph.review,read:2012.sph.w.dissipation.switches,abel:2011.sph.pressure.gradient.est}). All of these regimes are inevitable in most astrophysically interesting problems. Recently however, \citet{saitoh:2012.dens.indep.sph} (henceforth \SM) pointed out that the essential results of most of these flavors can be derived self-consistently in a manner that does properly conserve energy. The key insight is that the ``problematic'' inclusion of the density in the EOM (as opposed to some continuous property near contact discontinuities) arises because of the ultimately arbitrary choice of how to discretize the SPH {\em volume} element (typically chosen to be $\sim m_{i}/\rho_{i}$). Beginning with an alterative choice of volume element, one can in fact consistently derive a conservative EOM. They propose a specific form of the volume element involving internal energy and pressure, and show that this eliminates the surface tension term and resolves many problems of mixing near contact discontinuities. In this paper, we develop this approach to provide a rigorous, conservative, Lagrangian basis for the formulation of alternative ``flavors'' of SPH, and show that this can robustly resolve certain issues in mixing. Although the EOM derived in \SM\ conserves energy, it was derived from an ad-hoc discretization of the hydrodynamic equations, not the discrete particle Lagrangian. As such it cannot guarantee {\em simultaneous} conservation of energy and entropy (as well as momentum and angular momentum). And the EOM they derive is conservative only for constant SPH smoothing lengths (in time and space); to allow for adaptive smoothing (another major motivation for SPH), it is necessary to derive the ``$\nabla h$'' terms which account for their variations. This links the volume elements used for smoothing in a manner that necessitates a Lagrangian derivation. And their derivation depends on explicitly evolving the particle internal energy; there are a number of advantages to adopting entropy-based formulations of the SPH equations instead. We show here that -- allowing for a different initial choice of which thermodynamic volume variable is discretized -- an entire extensible class of SPH algorithms can be derived from the discrete particle Lagrangian, and write a general EOM for these methods (Eq.~\ref{eqn:eom}, our key result). We derive specific ``pressure-energy'' (Eq.~\ref{eqn:eom.pressure.energy.betterh}) and ``pressure-entropy'' (Eq.~\ref{eqn:eom.pressure.entropy.betterh}) formulations of the EOM, motivated by the approaches above that endeavor to enforce single-valued SPH pressures near contact discontinuities. We consider these methods in a wide range of idealized and more complex test problems, and show that they {\em simultaneously} maintain manifest conservation while tremendously improving the treatment of contact discontinuities and fluid mixing processes. \vspace{-0.5cm}
|
\label{sec:discussion} With inspiration from \SM, we re-derive a fully self-consistent set of SPH equations of motion which are independent of the kernel-calculated density, and therefore remove the well-known ``surface tension'' terms that can suppress fluid mixing. The equations still depend on the medium having a differentiable pressure (hence still require some artificial viscosity to capture shocks), but unlike the traditional SPH EOM, remain valid through contact discontinuities. Our derivation of the EOM relies on the key conceptual point in \SM, that the SPH volume element does not have to explicitly involve the mass density. However, our derivation and resulting EOM adds four key improvements: {\bf (1)} We rigorously derive the equations from the discretized particle Lagrangian. This guarantees one of the most powerful features of SPH, namely {\em manifest} simultaneous conservation of energy, entropy, momentum, and angular momentum, and an exact solution to the particle continuity equation. {\bf (2)} We similarly derive the ``$\nabla h$'' terms, which are {\em required} for manifest conservation if the SPH smoothing lengths $h_{i}$ are not everywhere constant. {\bf (3)} We derive an ``entropy formulation'' of the equations that allows for the direct evolution of the entropy, avoiding the need to construct/evolve an energy equation, and gives better entropy conservation properties as in \Sentropy; this also happens to minimize the correction terms involved in using a ``particle neighbor number'' definition to define $h$, as compared to the ``energy formulation.'' {\bf (4)} We show how the Lagrangian derivation can be generalized to {\em separate} definitions of the thermodynamic volume element (relating e.g.\ $P$ and $u$) and that used to define the smoothing lengths. This resolves problems of numerical stability and excess diffusion in strong shocks and/or large density contrasts, and automatically allows for varying particle masses. In fact, we derive a completely general, Lagrangian form of the EOM, including the $\nabla h$ terms, for {\em any} definition of the SPH thermodynamic volume element. Essentially {\em any} particle-carried quantity can be used in the kernel sum entering the EOM, and any (not necessarily the same) differentiable function used to define how the smoothing lengths $h_{i}$ are scaled. In some ways this replaces the long-known ``free weighting functions'' used to define the SPH EOM in their original ``discretized volume element'' formulation. However, in that approach, the choice of different functions generically violates conservation and continuity; here, we demonstrate that a similar physical degree of freedom can be utilized in the discretization of the equations of motion without such violations. Based on this degree of freedom, it is easy to see how different discretizations of the EOM might be optimized for some problems. By choosing the required kernel-evaluated element to directly represent a very smooth/stable property in the system, one not only removes spurious ``tension'' terms associated with discontinuities in other system variables, but also minimizes the inevitable discretization error from representing these quantities with a kernel sum. For the constant-pressure (but mixed density) fluid mixing tests we show here, the optimal choice is the ``pressure formulation.'' In MHD applications this kernel sum could trivially be altered to include the magnetic pressure. However if simulating an incompressible or weakly compressible fluid, the ``density formulation'' may well be superior. Direct kernel sums of nominally ``higher order'' properties such as the vorticity or vortensity are also valid and may represent useful formulations for some problems. It is even possible (in principle) to generalize our derivation to one in which different particle subsets have differently-defined volume elements; although we caution that such an approach requires great care. For the test problems here, we show that the ``pressure-entropy'' and ``pressure-energy'' formulations dramatically improve the treatment of fluid interface instabilities including the Kelvin-Helmholtz instability, Rayleigh-Taylor instability, and the ``blob test'' (a mix of Kelvin-Helmholtz and Raleigh-Taylor instabilities as well including non-linear evolution and shock capturing); giving results very similar to grid methods. They also remove the ``deforming'' effect of the surface tension term (allowing, for example, the long-term evolution of an irregular shape of gas at constant pressure but high density contrast); deformation is difficult to avoid even in grid codes (unless the chosen geometry matches the grid), and would otherwise require moving-mesh approaches to follow. However, unlike some of the modifications in the literature proposed to improve the fluid mixing in SPH (which violate conservation), the manifest conservation properties of our derivation mean that it remains well-behaved even in very strong shocks and does not encounter problems of either energy conservation or particle order in e.g.\ extremely strong blastwave problems. With these changes in place, we find weaker (albeit still significant) residual effects from improvement in the artificial viscosity scheme. Comparisons of such schemes are well-studied and what we implement here is still not the most sophisticated possible treatment, although it still considerably reduces artificial viscosity away from shocks \citep[for more detailed studies, see e.g.][]{cullen:2010.inviscid.sph}. We find similar effects from changes to the SPH smoothing kernel. Our favored kernel is taken from more detailed kernel comparison studies in \citet{hongbin.xin:05.sph.kernels,dehnen.aly:2012.sph.kernels}; however, unlike some other SPH formulations, we find that even the ``simplest'' kernel possible ($\NNb=32$ cubic spline) reproduces good results in several tests, except in the expected regime where we wish to resolve kernel-scale growing instabilities that rely on sub-sonic motions at the level of $\mathcal{M} \lesssim N_{\rm NGB}^{-1}$ and so summation errors dominate. This relates to the manifest conservation and maintenance of good particle order implicit in the EOM \citep[see][]{price:2012.sph.review}. We test the algorithm not just in the ``standard'' set of test problems but also an example of direct astrophysical interest, simulating the evolution of galaxies with a multi-phase ISM. This is useful because it makes clear that for this problem, at least, the differences arising from the treatment of different physics (e.g.\ how cooling, star formation, stellar feedback, and AGN feedback are implemented) makes, on average, larger differences than the numerical scheme \citep[also shown in other code comparisons; e.g.][]{scannapieco:2012.aquila.cosmo.sim.compare}. This is not surprising: those choices lead to orders-of-magnitude differences as opposed to the (still significant) factor $\sim$couple effects of numerical choices. Moreover the differences we are concerned with here largely pertain to mixing in sub-sonic, non-radiative flows dominated by thermal pressure; in contrast many astrophysical problems of interest involve highly super-sonic, radiative, gravity-dominated flows. In that limit, the differences owing to the algorithm are often -- though certainly not always -- minimized (see references in \S~\ref{sec:intro}). But there are important regimes with transonic flows where the numerical approach can make larger differences \citep[e.g.\ cosmological inflows \&\ outflows; see][]{vogelsberger:2011.arepo.vs.gadget.cosmo,keres:2011.arepo.gadget.disk.angmom,torrey:2011.arepo.disks}. Even in idealized test problems, we caution that simple physical differences can produce larger distinctions than the numerical method. For example, for several fluid mixing problems considered here, the ``correct'' MHD solution in the presence of an equipartition magnetic field can resemble the ``standard'' (density-entropy) SPH solution without a magnetic field, as opposed to the results from our pressure-entropy formulation or grid codes without such fields. The reason is that the real magnetic tension suppresses mixing, similar to the (purely numerical) ``surface tension'' term discussed in the text (so one might obtain a more ``realistic'' solution, but for entirely wrong reasons). Ultimately, the numerical formulations derived here should provide the basis for a more rigorous approach to the ``flavors'' of SPH, and a means to compare the consequences of the fundamental choice of how to discretize any SPH approach. This change to the algorithm is not a panacea! Fortunately, the modified equations of motion proposed here can be trivially incorporated with many other methods that improve on other numerical aspects, for example the inviscid algorithm in \citet{cullen:2010.inviscid.sph}, the higher-order dissipation switches in \citet{price:2008.sph.contact.discontinuities,rosswog:2010.relativistic.sph} and \citet{read:2012.sph.w.dissipation.switches}, and/or the gradient error reducing integral formulation of the kernel equations in \citet{garciasenz:2012.integral.sph}. We wish to stress that -- although SPH certainly has some disadvantages which we have not attempted to address here -- poor fluid mixing in contact discontinuities is not necessarily an ``inherent'' property of SPH. This problem can be improved without requiring additional dissipation terms (and without additional computational expense) while retaining what is probably the greatest advantage of SPH algorithms, namely their excellent conservation properties. \vspace{-0.7cm}
| 12
| 6
|
1206.5006
|
1206
|
1206.4146_arXiv.txt
|
{Rotational mixing in massive main sequence stars is predicted to monotonically increase their surface nitrogen abundance with time.} {We use this effect to design a method for constraining the age and the inclination angle of massive main sequence stars, given their observed luminosity, effective temperature, projected rotational velocity and surface nitrogen abundance.} {This method relies on stellar evolution models for different metallicities, masses and rotation rates. We use the population synthesis code STARMAKER to show the range of applicability of our method.} {We apply this method to 79 early B-type main sequence stars near the LMC clusters NGC 2004 and N 11 and the SMC clusters NGC 330 and NGC 346. From all stars within the sample, 17 were found to be suitable for an age analysis. For ten of them, which are rapidly rotating stars without a strong nitrogen enhancement, it has been previously concluded that they did not evolve as rotationally mixed single stars. This is confirmed by our analysis, which flags the age of these objects as highly discrepant with their isochrone ages. For the other seven stars, their nitrogen and isochrone ages are found to agree within error bars, what validates our method. Constraints on the inclination angle have been derived for the other 62 stars,with the implication that the nitrogen abundances of the SMC stars, for which mostly only upper limits are known, fall on average significantly below those limits.} {Nitrogen chronology is found to be a new useful tool for testing stellar evolution and to constrain fundamental properties of massive main sequence stars. A web version of this tool is provided.}
|
\label{sec1} The age determination of stars is an important objective, and several methods are currently being used. For star clusters, which presumably contain stars of similar ages, comparing the main sequence turn-off point in a color-magnitude or Hertzsprung-Russell diagram with stellar evolution models can predict the age of all stars in the cluster \citep{Jia_2002,Buonanno_1986}. For individual stars, ages can be derived by comparing their location in the HR-diagram with isochrones computed from stellar evolution models. In the envelope of low mass main sequence stars lithium is depleted, resulting in a monotonous decrease of the surface lithium abundance with time, which can be used to determine their age \citep{Soderblom_2010,Randich_2009}. Massive stars fuse hydrogen to helium via the CNO cycle. The reaction $\element[][14]{N} \left(\mathrm{p},\gamma\right) \element[][15]{O}$ has the lowest cross section, resulting in an increase of the nitrogen abundance compared to carbon. The evolution of massive stars is furthermore thought to be affected by rotation, which may cause mixing processes including large scale circulations and local hydrodynamic instabilities \citep{Endal_1987}. The nitrogen produced in the core can be transported to the surface when rotational mixing is considered, causing a monotonic increase of the surface nitrogen abundance with time \citep{Heger_2000,Maeder_2000a,Przybilla_2010}. Here, we suggest to utilize this effect to constrain the ages of stars. Based on detailed stellar evolution models, we present a method which reproduces the surface nitrogen abundances of massive main sequence stars as a function of their age, mass, surface rotational velocity and metallicity. Three stellar evolution grids for the Milky Way (MW), the Large Magellanic Cloud (LMC) and the Small Magellanic Cloud (SMC) are used, providing stellar models of different metallicities, masses and rotational velocities and include the effects of rotational mixing \citep{Brott_2010a}. The mixing processes in these stellar evolution models were calibrated to reproduce the massive main sequence stars observed within the VLT-FLAMES Survey of Massive Stars \citep{Evans_2005,Evans_2006}. The abundance analysis of massive stars is often focused on stars with small $v \sin i$ values, as this allows to obtain abundances with higher accuracies \citep{Przybilla_2010}. However, those objects are not well suited for nitrogen chronology, as they are either slow rotators, or have a low inclination angle. In the latter case, the age can not be well constrained, since a large range of true rotational velocities is possible. Many stars analyzed in the VLT-FLAMES Survey, however, are particularly suited as targets for nitrogen chronology, since this survey did not concentrate on the apparent slow rotators. On the other hand, results from the VLT-FLAMES Survey identified two groups of early B-type stars which could not be reconciled by the evolutionary models of single stars with rotational mixing \citep{Hunter_2008,Brott_2010b}. The first group consists of slowly rotating stars which are nitrogen rich, to which the method of nitrogen chronology is not applicable. The second group contains relatively fast rotating stars with only a modest surface nitrogen enhancement. While such stars are in principle predicted by models of rotating stars \citep{Brott_2010a}, they were not expected in be found in the VLT-FLAMES Survey due to the imposed V-band magnitude cut-off, since unevolved main sequence stars are visually faint, and the visually brighter more evolved rapid rotators are expected to be strongly nitrogen enriched \citep{Brott_2010b}. \citet{Brott_2010b} suggested that either these stars are the product of close binary evolution, or that rotational mixing is much less efficient than predicted by the stellar evolution models. The nitrogen chronology is able to shed new light on this issue. In contrast to the population synthesis pursued by \citep{Brott_2010b}, it does not use statistical arguments but is applied to each star individually. Its results can therefore be easily compared with classical age analyses, e.g., using isochrone fitting in the HR diagram. In this paper, we apply the nitrogen chronology to all LMC and SMC stars from the VLT-FLAMES Survey which have a projected rotational velocity larger than 100\,km/s. We organize our paper as follows. In Sect.~\ref{sec2} our method of nitrogen chronology is presented. Sect.~\ref{sec3} describes our analysis of the VLT-FLAMES sample. A discussion and evaluation of our new method is given in Sect.~\ref{sec4} and \ref{sec5}. Sect.~\ref{sec6} will round off with our conclusions.
|
\label{sec6} We present a new method of nitrogen chronology, which can constrain the age of a star, and/or its inclination angle, based on its observed surface nitrogen abundance, mass and projected surface rotational velocity, by comparing the observed nitrogen abundance with the one predicted by the theory of rotational mixing in single stars. This method can be applied to stars when their surface nitrogen abundance increases monotonically with time during the main sequence evolution. It has been worked out here for stars in the mass range between 5\,$M_{\odot}$ and up to 35--50\,$M_{\odot}$, and for metallicities adequate for the Milky Way, the LMC and the SMC, based on the stellar evolution models of \citet{Brott_2010a}. We apply our method to 79 stars from the early B~type LMC and SMC samples of the VLT-FLAMES Survey of Massive Stars \citep{Evans_2005,Evans_2006}. Age constraints from nitrogen chronology could be obtained for 17~of the 79 analyzed stars. In Sect.~4.2, we compared those to ages of these stars as derived from isochrone fitting in the HR diagram (Fig.~\ref{fig_comparison}). We found the isochrone ages of 10 objects to be incompatible with their nitrogen age constraints. Based on their rapid rotation and low nitrogen enrichment, \citet{Brott_2010b} concluded that these 10~stars did not evolve as rotationally mixed single stars. Our star-by-star analysis of these objects is reinforcing their conclusion that the weakly enriched fast rotators in the LMC early B~stars of the FLAMES Survey --- which are 15\% of the survey stars --- can not be explained by rotating single stars. For the remaining 7~stars with nitrogen age constraints, six LMC and one SMC star, nitrogen and isochrone ages are found to be consistent within error estimates. Four of them show good agreement, which also allows to derive their inclination angles (Sect.~4.2). We conclude that our nitrogen chronology results are consistent with the isochrone method for stars which have no indication of an unusual evolution, but incompatible for stars which are suspected binary product according to \citet{Brott_2010b}. Therefore, our new method provides important new results for both groups of stars. However, it is clearly desirable to obtain larger samples of stars especially of the first group. For the 62~stars (28 LMC and 34 SMC stars) for which no age constraint could be derived, our method provides limits on the inclination for each object (Sect.~5). While the results for the LMC stars are roughly consistent with random inclination angles, this appears to be different for the SMC stars, for which only upper limits to the nitrogen abundance are available in most cases \citep{Hunter_2009}. We argue that random inclinations are also realized for the SMC stars, with the implication that the true nitrogen surface abundances of most SMC stars are significantly below the derived upper limits. We believe that our new method can be used in the near future to provide further tests of the theory of rotational mixing in stars, and --- if such tests converge to confirm this theory --- to act as a new chronometer to constrain the ages of massive main sequence stars. Based on the method presented here, a web tool\footnote{http://www.astro.uni-bonn.de/stars/resources.html} can be found online. For given mass, (projected) surface rotational velocity and metallicity, the surface nitrogen abundance is calculated and compared to the observed surface nitrogen abundance to constrain the age and inclination angle.
| 12
| 6
|
1206.4146
|
1206
|
1206.4693_arXiv.txt
|
We predict the space density of molecular gas reservoirs in the Universe, and place a lower limit on the number counts of carbon monoxide (CO), hydrogen cyanide (HCN) molecular and \cii atomic emission lines in blind redshift surveys in the submillimeter--centimeter spectral regime. Our model uses:\ (a) recently available HCN Spectral Line Energy Distributions (SLEDs) of local Luminous Infrared Galaxies (LIRGs, $L_{\rm IR}>10^{11}L_\odot$), (b)\ a value for $\epsilon_{\star}$=$\rm SFR$/$M_{\rm dense}(\rm H_2)$ provided by new developments in the study of star formation feedback on the interstellar medium and (c)\ a model for the evolution of the infrared luminosity density. Minimal `emergent' CO SLEDs from the dense gas reservoirs expected in all star-forming systems in the Universe are then computed from the HCN SLEDs since warm, HCN-bright gas will necessarily be CO-bright, with the dense star-forming gas phase setting an obvious minimum to the total molecular gas mass of any star-forming galaxy. We include \cii as the most important of the far-infrared cooling lines. Optimal blind surveys with the Atacama Large Millimeter Array (ALMA) could potentially detect very distant ($z\sim10$--$12$) \cii emitters in the $\geq$ULIRG galaxy class at a rate of $\sim$0.1--1 per hour (although this prediction is strongly dependent on the star formation and enrichment history at this early epoch), whereas the (high-frequency) Square Kilometer Array (SKA) will be capable of blindly detecting $z>3$ low-{\it J} CO emitters at a rate of $\sim$40--70 per hour. The \cii line holds special promise for the detection of metal-poor systems with extensive reservoirs of CO-dark molecular gas where detection rates with ALMA can reach up to 2--7 per hour in Bands 4--6.
|
Since the first detections of the $J$$=$$1$$\rightarrow$$0$ rotational transition of $^{12}$CO and some of its isotopologues in Galactic molecular clouds ($^{13}$CO, C$^{18}$O) (Wilson et al.\ 1970; Penzias et al.\ 1971, 1972), and in galactic nuclei (Rickard et al.\ 1975) there have been many studies of CO line emission in galaxies using single dish radio telescopes and interferometer arrays (for reviews see Young \& Scoville\ 1991 and Solomon \& Vanden Bout\ 2005). Multi-{\it J} CO line ratio surveys are now routinely used to assess the state of the molecular gas in galaxies (e.g.\ Braine \& Combes\ 1992; Aalto et al.\ 1995; Papadopoulos \& Seaquist\ 1998; Mauersberger et al.\ 1999; Nieten et al.\ 1999; Yao et al.\ 2003; Mao et al.\ 2011) over the density regime where most of its mass resides ($n\sim 10^{2-3}\,{\rm cm}^{-3}$). The fainter molecular line emission from heavy-rotor molecules such as HCN have also become a standard tool for assesing the state and the mass of the denser ($>$$10^4\,{\rm cm}^{-3}$) gas phase where stars actually form in Giant Molecular Clouds (GMCs, e.g.\ Nguyen-Q-Rieu et al.\ 1989; Solomon et al.\ 1992a; Paglione et al.\ 1995, 1997; Jackson et al.\ 1995). The role of the latter phase as the direct fuel of star formation in individual GMCs, quiescent disks and merger-driven spectacular starbursts in the local and distant Universe is now well established over an astounding 7--8 orders of magnitude (Gao \& Solomon\ 2004; Wu et al.\ 2005; Juneau et al.\ 2009; Wang et al.\ 2011). In the past decade, numerous high-{\it z} detections have revealed the fundamental role of molecular lines in assessing the state and mass of the molecular gas, and the dynamical mass of heavily dust-enshrouded galaxies in the early Universe (Solomon \& Vanden Bout 2005\ and references therein), and in some cases have provided remarkable insights into the properties of the molecular interstellar medium (ISM) in early galaxies (see Danielson et al.\ 2010 for a recent example of a well sampled CO spectral line energy distribution (SLED) in a $z=2.3$ gravitationally lensed galaxy). This exploration began with the first detection of \coiii\ line emission in the strongly-lensed distant dust-enshrouded galaxy IRAS\,10214+4724 at $z\sim 2.3$ (Brown \& Vanden Bout\ 1991; Solomon et al.\ 1992b). It continued with the detection of CO transitions in distant submm-bright galaxies (SMGs, $L_{\rm IR}>10^{12}L_\odot$) (Frayer et al.\ 1998, 1999; Greve et al.\ 2005) and is increasingly encompassing less extreme, but still massive systems such as Lyman Break galaxies (Baker et al.\ 2004), optical/near-infrared selected galaxies at $z\sim1.5$ (Dannerbauer et al.\ 2009; Daddi et al.\ 2010) and fortuitously lensed systems (Danielson et al.\ 2011; Lupu et al.\ 2011). Several spectacular CO line detections have also been obtained also for other high-redshift systems such as radio galaxies (e.g.\ De Breuck et al.\ 2005) and QSOs out to $z\sim 6.4$. This epoch is close to the era of reionization -- the final frontier of galaxy evolution studies -- revealing the gas-rich hosts to rapid galaxy growth at these early times (Walter et al.\ 2003, 2004; Weiss et al.\ 2007). These discoveries and advancements, made possible as sensitivities of millimeter/submillimeter interferometer arrays improved in the last decade, still yield only a glimpse of what will be a new era where molecular and atomic (e.g.\ \cii\,$\lambda158$, \ci) ISM lines will replace nebular optical/near-infrared (OIR) lines as the main tool of choice for discerning galaxy formation and evolution across the full span of cosmic time pertinent to galaxy growth, from the end of the reionisation epoch ($z>7$) to the present (Walter \& Carilli\ 2008). Direct `blind' searches of gas-rich galaxies using submm--cm wave molecular and atomic lines are the only tool that can:\ (i) uniformly select galaxies according to their molecular gas content rather than their SFR and the star formation efficiency (a bias that has so far -- necessarily -- affected all high-$z$ gas studies),\ (ii) immediately provides redshifts and eventually dynamical mass information,\ (iii) holds the promise of discovering large outliers of the local $\Sigma _{\rm SFR}$-$\Sigma_{{\rm H}_2}$ (Schmidt--Kennicutt) relations (Kennicutt\ 1998), with large reservoirs of molecular gas but low levels of SFR (Papadopoulos \& Pelupessy\ 2010), and\ (iv) can possibly determine the star formation `mode' (starburst/merger-driven versus quiescent/disk-like), in a uniform and extinction-free manner (Paper II, Papadopoulos \& Geach\ 2012). Well-sampled CO spectral line energy distributions, and their robust normalization by some observable galaxy property are necessary for predicting the emergent CO line luminosities in star-forming systems. The lack of these two key ingredients translates to major uncertainties for the source counts predicted for blank-field cm/mm/submm molecular line surveys (Combes et al.\ 1999; Blain et al.\ 2000; Carilli \& Blain\ 2002), as well as the frequency and flux range where such surveys become optimal (Blain et al.\ 2000). The dense gas phase ($n({\rm H}_2)$$>$10$^{4}$\,cm$^{-3}$) and the linear relation of its mass to the SFR in individual GMCs ($L_{\rm IR}$$\sim $10$^{4.5}$\,$L_{\odot}$), ultraluminous infrared galaxies (ULIRGs, $L_{\rm IR}$$\sim $$10^{12}$\,$L_{\odot}$) and high-redshift extreme starbursts (HLIRGs, $L_{\rm IR}$$\sim $$10^{13}$\,$L_{\odot}$) makes it an obvious benchmark for computing minimum emergent molecular line luminosities. Indeed only this gas phase yields physically meaningful estimates of the so-called star formation efficiency (and its equivalent interpretation in terms of gas consumption timescales) while the HCN-deduced (and thus well-excited CO SLEDs) contain minimal uncertainties up to high-{\it J} rotational transitions of CO. The dense, HCN-bright, molecular gas phase in galaxies is thus an obvious ingredient of any theoretical models for blank-field cm/mm/submm molecular line surveys.
|
We have presented a conservative model of the number counts of galaxies detected in a blind molecular line survey in the sub-mm/mm/cm regime. Our model calculates the `emergent' CO, HCN and \cii $\lambda$158$\mu$m emission of star-forming galaxies, and is rooted in the latest models of star formation feedback and empirical data on the HCN SLED (tracing the dense gas phase) in local star-forming galaxies. The normalization of the emergent CO SLED is given by the star formation rate, which in this case is taken to be the infrared luminosity of a galaxy. Thus, our model describes the {\it minimum} molecular line emission expected for star-forming galaxies based solely on the luminosity of their actively star-forming reservoirs. This could be used to design follow-up spectroscopic surveys for an unbiased $L_{\rm IR}$ limited survey. Coupled with an up-to-date model for the evolution of the infrared luminosity density that successfully re-produces the observed number counts of galaxies over a wide range of the infrared wavebands (B\'ethermin et al.\ 2011), we make predictions of the lower limit of integrated number counts of line-emitting galaxies across a range of observed frequencies and bandpasses pertinent to the main facilities capable of performing a molecular redshift survey (ALMA, SKA and its pathfinders). We consider ambitious blind redshift surveys, working at the optimal flux limit set by the predicted knee in the galaxy number counts, and discarding information about the shape of the spectral line (i.e.\ binning to a spectral resolution of 1000, i.e.\ $\sim$300\,km\,s$^{-1}$). Such blind surveys can reveal insight into: \medskip {\it The epoch of re-ionization:}\ The sensitive ALMA bands could potentially detect ULIRG-class \cii emitters close to the epoch of re-ionisation, $z\gtrsim 10$, at a rate of up to one per hour (although this is highly sensitive to the star formation history of the Universe at this early time). Nevertheless, should such extreme systems exist at this epoch, a blind ALMA survey would be capable of finding them, and their abundance would provide valuable insight into the star formation and chemical history of the Universe close to the era when the first stars ignited. In our minimal model, \cii emitters dominate blind (optimal) surveys with ALMA, however mid-{\it J} CO emitters would also be detected at lower rates, but with increasing yields for deeper (but sub-optimal) surveys. \medskip {\it CO-dark galaxies:}\ We also examine the possibility of detecting \cii luminous, but CO-dark gas reservoirs in metal-poor galaxies at high-{\it z} with ALMA. Assuming such a population exists with a similar space density to Lyman Break Galaxies, blind surveys with ALMA could detect systems at $z\sim8$--$12$ with optimal rates of $\sim$2--7 per hour. \medskip {\it Efficient blind surveys of low-{\it J} CO emitters at $z\gtrsim3$:}\, The SKA will represent a sea-change in the sensitivity of radio/cm-wave surveys, with SKA Phase 3 (offering access to the radio {\it K} band) providing access to low-{\it J} CO emission at $z>3$. We predict that an optimal redshift survey could detect $\sim$30--70 ULIRG-class CO emitters per hour. While our model is based on the abundance of star-forming galaxies, blind SKA surveys could also detect outliers from the standard Schmidt-Kennicutt relation. In a follow-up work, Paper\ II (Papadopoulos \& Geach 2012), we consider the detectability of `pre-starburst' galaxies, representing a brief gas-rich phase preceding the onset of an episode of intense star formation where the host galaxy is extremely difficult to detect in any other waveband. \medskip \noindent The coming decade and the years beyond will be an exciting time for extragalactic astronomy: we will routinely detect molecular emission from high-redshift galaxies, breaking through the sensitivity floor that has limited the majority of current studies to the most luminous or fortuitously gravitationally lensed galaxies. This work presents a simple, empirically-based model to aid in the design of redshift surveys (both blind and semi-blind). Although we promote ambitious observations, with -- arguably -- speculative results, we are motivated by the rich spoils: totally new and, in some cases, unique insights into the physics of galaxy formation that could be the reward for such efforts.
| 12
| 6
|
1206.4693
|
1206
|
1206.2140_arXiv.txt
|
{} {Deuterated molecules have been detected and studied toward Orion BN/KL in the past decades, mostly with single-dish telescopes. However, high angular resolution data are critical not only for interpreting the spatial distribution of the deuteration ratio but also for understanding this complex region in terms of cloud evolution involving star-forming activities and stellar feedbacks. Therefore, it is important to investigate the deuterated ratio of methanol, one of the most abundant grain-surface species, on a scale of a few arcseconds to better understand the physical conditions related to deuteration in Orion BN/KL.} {Orion BN/KL was extensively observed with the IRAM Plateau de Bure Interferometer from 1999 to 2007 in the 1 to 3 mm range. The angular resolution varies from $1\farcs8\times0\farcs8$ to $3\farcs6\times2\farcs3$ and the spectral resolution varies from 0.4 to 1.9 \kms. Deuterated methanol \dmeth\ and \methd\ and \meth\ lines were searched for within our 3 mm and 1.3 mm data sets.} {We present here the first high angular resolution ($1\farcs8\times0\farcs8$) images of deuterated methanol \dmeth\ in Orion BN/KL. Six \dmeth\ lines were detected around 105.8, 223.5, and 225.9 GHz. In addition, three E-type methanol lines around 101--102 GHz were detected and were used to derive the corresponding \meth\ rotational temperatures and column densities toward different regions across Orion BN/KL. The strongest \dmeth\ and \meth\ emissions come from the Hot Core southwest region with a velocity that is typical of the Compact Ridge (\vlsr\ $\approx8$ \kms). We derive [\dmeth]/[\meth] abundance ratios of $0.8-1.3\times10^{-3}$ toward three \dmeth\ emission peaks. A new transition of \methd\ was detected at 226.2 GHz for the first time in the interstellar medium. Its distribution is similar to that of \dmeth. Besides, we find that the [\dmeth]/[\methd] abundance ratios are lower than unity in the central part of BN/KL. Furthermore, the HDO $3_{1,2}-2_{2,1}$ line at 225.9 GHz was detected and its emission distribution shows a shift of a few arcseconds with respect to the deuterated methanol emission that likely results from different excitation effects. The deuteration ratios derived along Orion BN/KL are not markedly different from one clump to another. However, various processes such as slow heating due to ongoing star formation, heating by luminous infrared sources, or heating by shocks could be competing to explain some local differences observed for these ratios.} {}
|
\begin{table*} \caption{Observational parameters of the PdBI data sets} % \label{table-data} % \centering % \begin{tabular}{lcccccccc} % \hline\hline % Bandwidth & Observation date & Configuration & Flux conversion & RMS noise & $\theta_{\rm HPBW}$\tablefootmark{a} & $\delta{\rm v}$\tablefootmark{b} & \multicolumn{2}{c}{$\theta_{\rm syn}$\tablefootmark{c}} \\ (GHz) & & & (1 \jb) & (\mjb) & (\arcsec) & (\kms) & (\arcsec$\times$\arcsec) & PA (\degr) \\ \hline % 101.178--101.717 & 2003--2006 & BC & 15.8 K & 2.9 & 49.7 & 1.85 & $3.79\times1.99$ & 22 \\ 105.655--105.726 & 2005 Aug--Nov & D & 2.9 K & 7.6 & 47.7 & 0.89 & $7.13\times5.36$ & 9 \\ 223.402--223.941 & 2003--2007 & BC & 17.3 K & 21.2 & 22.5 & 0.84 & $1.79\times0.79$ & 14 \\ 225.805--225.942 & 2005 Sep--Nov & D & 2.9 K & 40.8 & 22.3 & 0.42 & $3.63\times2.26$ & 12 \\ 225.990--226.192 & 2005 Sep--Nov & D & 2.9 K & 40.8 & 22.3 & 0.42 & $3.63\times2.25$ & 12 \\ \hline % \end{tabular} \tablefoot{ \tablefoottext{a}{Primary beam size} \tablefoottext{b}{Channel separation} \tablefoottext{c}{Synthesized beam size} } \end{table*} Deuterium chemistry in the interstellar medium (ISM) has been intensively studied in recent decades. More and more deuterated molecules have been found as well as multiply-deuterated species, e.g., ND$_{3}$ \citep{van der Tak2002,Lis2002} and CD$_3$OH \citep{Parise2004}. Deuterium chemistry models have also been revised to explain the observations that show a strong enhancement of deuterium-bearing molecule abundances in star-forming regions \citep[e.g.,][]{Roberts2003,Roberts2000,Charnley1997}, compared with the D/H ratio of $2-3\times10^{-5}$ in the local interstellar gas \citep[see, e.g.,][and references therein]{Linsky2006}. Those strong enhancements (by a factor of a few thousand) are seen mostly in methanol, ammonia, water, and formaldehyde, leading to the abundance ratios $>0.1$ compared with their non-deuterated analogs. The formation of these molecules can be largely explained by grain-surface reactions that provide a natural explanation for the abundant doubly or multiply deuterated molecules that ion-molecule chemistry failed to predict \citep{Turner1990}. Many deuterated molecules were first detected toward the Orion Becklin-Neugebauer/Kleinmann-Low (BN/KL) region \citep{Becklin1967,Kleinmann1967}, one of the closest \citep[$414\pm7$ pc,][]{Menten2007} and most-studied star-forming regions in the sky. For example, deuterated water HDO and deuterated ammonia NH$_2$D were first detected by \citet{Turner1975} and \citet{Rodriguez1978}, and the single-deuterated methanol molecules \methd\ and \dmeth\ were first detected toward the same region by \citet{Mauersberger1988} and \citet{Jacq1993}, respectively. Although grain-surface chemical models can well explain the high abundance of those deuterated molecules in molecular clouds, the deuteration branching ratios in the same species predicted by models disagree with observations. For instance, \citet{Charnley1997} showed that the formation of deuterated methanol on grains based on the addition of H and D atoms to CO always leads to [\dmeth]/[\methd] abundance ratios of about 3. This prediction is in conflict with the [\dmeth]/[\methd] ratio of 1.1--1.5 observed by \citet{Jacq1993} in Orion BN/KL. However, \citet{Rodgers2002} later pointed out that without taking into account other surface species such as CO and H$_2$CO and possibly different energy barriers involved in the reaction scheme, the [\dmeth]/[\methd] ratio of 3 is an artificial result of the model assumptions. In addition, most proposed models that include surface chemistry strongly depend on local environment where temperature and gas density play important roles in the chemical reaction rates. This is especially true for the Orion BN/KL region where contributions from stellar feedbacks (e.g., ultraviolet photons) and star formation activities (e.g., outflows/shocks) alter the warm-up history of the cloud, involving both grain surface and gas-phase chemistry. Hence, it is crucial to investigate the BN/KL deuterated methanol distribution with high spatial resolutions so that the physical conditions of individual clumps are properly constrained and the deuteration ratios can be related to specific physical processes. Additionally, deeper insight into the processes at play can be gained by comparing \dmeth\ maps with another major deuterated species, HDO. The massive star-forming region Orion BN/KL is complicated not only because of the interactions between outflows and the ambient material, but also because of its rich and complex chemistry at the so-called Orion Hot Core and Compact Ridge regions \citep[see, e.g.,][]{Blake1987}. Owing to larger single-dish beam sizes in the early spectral line studies of Orion BN/KL, molecular line profiles were usually decomposed into several components according to their local standard of rest (LSR) velocities and line widths. The Orion hot molecular core (Hot Core) is usually characterized by its velocity component at \vlsr=5--6 \kms\ and a broad line width of about 5--10 \kms. It has been identified in the interferometric maps as a strong mm/submm continuum emission peak close to the infrared (IR) source IRc2 \citep[see, e.g.,][]{Gezari1998,Downes1981,Rieke1973}, whereas many molecular emission peaks observed around 8 \kms\ are displaced from the Hot Core by about 4\arcsec\ to the southwest \citep[Hot Core Southwest, HC-SW; see, e.g., works of][and references therein]{Favre2011,Wang2010,Tang2010,Friedel2008}. The 8--9 \kms\ LSR velocity component of the Orion Compact Ridge exhibits a relatively narrow line width (3--5 \kms). The exact location of the Orion Compact Ridge is somewhat ambiguous, but recent interferometric observations \citep[e.g.,][]{Favre2011,Friedel2008} suggest that it is located 10\arcsec--15\arcsec\ to the southwest of the Hot Core (the strongest dust continuum peak). It is important to mention that the Orion Hot Core is located within the NE-SW dense ridge of the BN/KL region seen in dust continuum emission \citep[][]{Favre2011,Tang2010}, and is part of the hierarchical filamentary structure seen on a larger scale in OMC-1 \citep[e.g., mid-$J$ CO images by][]{Peng2012}. Additionally, the Orion Compact Ridge is located at the southern part of this dense ridge, the bottom part of the V-shaped region seen in many molecular lines, e.g., CS \citep{Chandler1997}, SO \citep{Wright1996}, HCOOCH$_3$ \citep{Favre2011}, and NH$_3$ \citep{Goddi2011}. The main goal of this paper is to investigate the deuteration ratios in Orion BN/KL, and address the possible causes for the abundant deuterated methanol in this region. We present the first high angular resolution ($1\farcs8\times0\farcs8$) images of \dmeth\ toward the Orion BN/KL region (\S\ref{dmeth-result}). \meth\ maps were also obtained from the same data sets with a $3\farcs8\times2\farcs0$ resolution (\S\ref{meth-result}). The \methd\ map and HDO result are shown in \S\ref{methd-result} and \S\ref{hdo-result}, respectively. We discuss methanol deuteration in \S\ref{meth-diss} for \dmeth\ and \S\ref{methd-diss} for \methd. In \S\ref{Herschel-compare}, our own \meth\ data are discussed in the light of the spectral line profiles obtained at much higher frequencies with {\it Herschel}. Comparison of our deuterated methanol and methanol maps with our deuterated water maps is presented in \S\ref{HDO-compare}, and water and methanol deuteration ratios are discussed in \S\ref{deuteration-ratio}.
|
The main findings and conclusions of our study based on observations of several transitions of deuterated methanol and one transition of deuterated water in Orion BN/KL are as follows. \begin{enumerate} \item We have obtained the first high angular resolution ($1\farcs8\times0\farcs8$) \dmeth\ images detected around 223.5 GHz toward Orion BN/KL and compared these data with somewhat lower resolution ($3\farcs8\times2\farcs0$) \meth\ images at 101.5 GHz. The strongest \dmeth\ and \meth\ emissions come from the Hot Core southwest region exhibiting an LSR velocity of about 8 \kms, typical of the Orion Compact Ridge region. The \dmeth\ emission is clumpy and the column densities are estimated to be about $1-9\times10^{15}$ \cmm\ toward these clumps. The \meth\ column densities are about $3-5\times10^{17}$ \cmm\ across Orion BN/KL, leading to a [\dmeth]/[\meth] deuteration ratio of $0.8-1.3\times10^{-3}$ toward three deuterated methanol clumps and below $2\times10^{-4}$ toward KL-W. \item The [\dmeth]/[\methd] abundance ratio map was obtained for Orion BN/KL, and their ratios are less than unity at the central part of the region. These ratios are lower than the statistical factor of 3 derived in the simplest deuteration models, and definitely lower than the values derived in low-mass protostars \citep[e.g.,][]{Parise2002,Ratajczak2011}. \item We have mapped with moderately high spatial resolution ($3\farcs6\times2\farcs3$) the 225.9 GHz transition of HDO and compared its distribution with \dmeth, \meth, and \thmeth. We find that the deuterated water ratio is about one order of magnitude lower than the deuterated methanol ratio. H-D substitution may explain why methanol is easier to be deuterated than water. \item The deuteration ratios derived in this work are not strongly different from one clump to another, except perhaps toward KL-W where more observations are desirable to conclude. However, to explain the slight differences observed locally in the abundance ratios of identified clumps, we suggest that various processes could be competing, for instance, heating by luminous infrared sources, or heating by shocks. \end{enumerate}
| 12
| 6
|
1206.2140
|
1206
|
1206.5626_arXiv.txt
|
We report results of a systematic study of the broad band (2--2000~keV) time resolved prompt emission spectra of a sample of gamma-ray bursts (GRBs) detected with both Wide Field Cameras on board the \sax\ satellite and the \batse\ experiment on board CGRO. In this first paper, we study the time-resolved dependence of the intrinsic peak energy $E_{p,i}$ of the $E F(E)$ spectrum on the corresponding isotropic bolometric luminosity $L_{\rm iso}$. The $E_{p,i}$--$L_{\rm iso}$ relation or the equivalent relation between $E_{p,i}$ and the bolometric released energy $E_{iso}$, derived using the time averaged spectra of long GRBs with known redshift, is well established, but its physical origin is still a subject of discussion. In addition, some authors maintain that these relations are the result of instrumental selection effects. We find that not only a relation between the measured peak energy $E_p$ and the corresponding energy flux, but also a strong $E_{p,i}$ versus $L_{\rm iso}$ correlation are found within each burst and merging together the time resolved data points from different GRBs. We do not expect significant instrumental selection effects that can affect the obtained results, apart from the fact that the GRBs in our sample are sufficiently bright to perform a time-resolved spectroscopy and that they have known redshift. If the fundamental physical process that gives rise to the GRB phenomenon does not depend on its brightness, we conclude that the found $E_{p,i}$ versus $L_{\rm iso}$ correlation within each GRB is intrinsic to the emission process, and that the correlations discovered by Amati et al. and Yonetoku et al. are likely not the result of selection effects. We also discuss the properties of the correlations found.
|
\label{intro} In spite of the major advances in knowledge of gamma--ray bursts (GRBs) afterglow properties, mainly made with \swift, the GRB phenomenon is still poorly understood \citep{Lyutikov09}. It is recognized that the study of the prompt emission is of crucial importance, as it is more directly connected with the original explosion. One of the issues still open is the radiation emission mechanism(s) at work. Most of the GRB properties derived thus far come from the time-averaged spectra. The function that has been found to better describe them from 15 keV up to 10 MeV is a smoothly broken power-law proposed by \citet[{\em Band function}, BF]{Band93}. On the basis of the spectral data obtained with the {\em Burst and Transient Source Experiment} (\batse), aboard the \gro\ satellite ({\em CGRO}) and, for example, with the \sax\ GRBM data \citep[e.g.,][]{Guidorzi11}, for long GRBs ($>$2 s), the mean value of the low-energy photon index $\alpha$ of the BF is about $-1$, while that of the high-energy photon index $\beta$ is about $-2.3$ \citep{Kaneko06}. As a consequence of this result, the received power per unit logarithmic energy interval $EF(E)$ shows a peak value that, in the \batse\ era, seemed to show a sharp Gaussian distribution around 200 keV. With the discovery of the X--ray flashes with \sax, later also found with {\em HETE}-2, \swift, and, now, with the \fermi\ Gamma-ray Burst Monitor, this distribution results in being much flatter \citep[e.g.,][]{Kippen03,Sakamoto05}. In the cases in which $\beta$ cannot be constrained, a power-law model with a high energy exponential cutoff ({CPL) gives a good fit to the data, and, in some cases, even a simple power-law can describe the GRB time averaged spectra up to several MeV photon energies \citep{Kaneko06}. Several radiative emission models have been worked out for the interpretation of the GRB spectra. Given their non-thermal shape, the first model proposed was synchrotron emission by non thermal electrons in strong magnetic fields \citep{Rees94,Katz94,Tavani96}. Indeed, the time-averaged spectra of many GRBs are consistent with an optically thin synchrotron shock model \citep[e.g.,][]{Tavani96,Amati01}. However, there are a significant number of GRBs for which this model does not work. Indeed, while for an optically thin synchrotron spectrum, the expected power-law index of the $E F(E)$ spectrum below the peak energy $E_p$ cannot be steeper than 4/3 (ideal case of an instantaneous spectrum in which electron cooling is not taken into account), in many cases \citep[e.g.,][]{Preece98,Preece00} the measured spectra, even those time resolved \citep{Crider97,Frontera00a}, are inconsistent with these expectations. To overcome these difficulties, either modifications of the above synchrotron scenario \citep[e.g.,][]{Lloyd00a} or other radiative models \citep[e.g.,][]{Liang97,Blinnikov99,Lazzati00,Meszaros00,Stern04,Peer06,Peer07,Lazzati09,Peer11} have been suggested. Each of these models interprets some of the prompt emission features, but fails to interpret others. One of the GRB spectral properties that the emission models should interpret is the correlation between the intrinsic (redshift corrected) peak energy $E_{p,i}$ of the $E F(E)$ function and either the GRB released energy $E_{{\rm iso}}$ \citep{Amati02} or the peak bolometric luminosity $L_{p,{\rm iso}}$ \citep{Yonetoku04}. Both correlations (the Amati relation and the Yonetoku relation) have been derived from the time-integrated spectra assuming isotropic emission. The Yonetoku relation followed the Amati's, from which the so-called Ghirlanda relation \citep{Ghirlanda04a} was also derived by replacing the released energy $E_{\rm iso}$ with that ($E_{\gamma}$) corrected for the beaming factor ($E_{\gamma} = (1-\cos{\theta}) E_{\rm iso}$). The latter is model dependent, being derived assuming a jet like structure of the fireball, a constant efficiency of the fireball in converting kinetic energy in the ejecta into gamma rays, a mass density distribution of the circumburst medium, and, mainly, that the break time observed in the late afterglow light curve occurs when reciprocal of the bulk Lorentz factor of the jet, $1/\Gamma$, becomes of the order of the jet opening angle $\theta_{\rm jet}$. The latter assumption requires that the break time is achromatic, property not observed in many \swift\ GRBs \citep[e.g.,][]{Campana07,Melandri08}. In any case, these relations are all equivalent as far as the physics: a relation between the photon energy in which most of the energy is released and the electromagnetic radiation emitted by a GRB, which can be expressed equivalently in terms of released energy or peak (or average) GRB luminosity. The Amati relation \citep[$E_{p,i} = K E_{\rm iso}^m$, with $K= 98\pm 7$ and $m = 0.54\pm0.03$, where $E_{p,i}$ is measured in keV and $E_{\rm iso}$ is given in units of $10^{52}$~erg]{Amati08}, is satisfied, within an extra-Poissonian scatter of $\log{E_{p,i}}$ normally distributed around the best-fit power-law with $\sigma\sim 0.2$~dex, by all long GRBs (more than 100) with known redshift $z$ discovered thus far, except the nearest and least energetic GRB ($z= 0.0085$) ever observed (GRB\,980425) and, maybe, GRB\,031203 \citep[e.g.][]{Amati07}, but not by short GRBs. Also the Yonetoku relation is satisfied for GRBs with known redshift, except GRB\,980425 (e.g., \citet{Ghirlanda05b,Nava12}), given the tight correlation found between $L_{p,{\rm iso}}$ and $E_{\rm iso}$ \citep{Ghirlanda05b}. In spite of this, the Amati relation has been questioned by various authors \citep{Band05,Butler07b,Butler09,Shahmoradi09,Collazzi11}, maintaining that it is likely the result of selection effects, even if these effects, when investigated by other authors \citep{Ghirlanda05a,Ghirlanda08,Nava08,Amati09,Krimm09,Nava11a} are found to be marginal. One of the major criticisms of the Amati relation is that its normalization depends on instrument sensitivity \citep{Butler07b}. However, this conclusion by \citet{Butler07b} for \swift\ GRBs was not based on measured spectral peak energies $E_{p,i}$, but on $E_{p,i}$ values inferred, under some assumptions, from a Bayesian method. An investigation, performed by \citet{Amati09} by deriving the $E_{p,i}$ dependence on $E_{\rm iso}$ for different sets of GRBs, each obtained with a different instrument, and by using the measured values of $E_{p,i}$ reported in the \swift\ BAT official catalog by \citet{Sakamoto08}, has not confirmed the inferences by \citet{Butler07b}. Other works that question the reliability of the $E_{p,i}$ -- $E_{\rm iso}$ make use of the ($E_p$, Fluence) observer plane, positioning in this plane the data points obtained from the spectral analysis of GRBs detected with different satellite instruments, using the method first proposed by \citet{Nakar05}. While some authors \citep[e.g.][]{Band05,Goldstein10} find that most of the long \batse\ GRBs do not satisfy the relation, other authors \citep[e.g.,][]{Ghirlanda05a,Nava08,Nava11a} find that most of them do. This discrepancy is due to various reasons, such as the condition assumed to satisfy or not satisfy the Amati relation and the systematic errors in the determination of fluence and $E_p$ \citep[e.g.][]{Collazzi11}. However some results are well established and are shared by most authors who have performed the $E_{p,i}$ -- $E_{\rm iso}$ test. GRBs with known redshift are not a special sub-population, they are evenly distributed along the weaker correlation between $E_p$ and fluence \citep{Ghirlanda08,Collazzi12}. No evidence of evolution of the $E_{p,i}$ -- $E_{\rm iso}$ correlation with the redshift is found \citep{Ghirlanda08,Nava11a}. In the observer frame, after all the results obtained with \swift\ and \fermi, no bursts with large fluence and low/intermediate $E_p$ have been found. They do not exist or should be very rare: Nothing prevents their detection. Instead, bursts with intermediate/high $E_p$ and small fluence could be affected by instrumental selection effects, such as the minimum flux to trigger a GRB, the minimum fluence to fit its spectrum and constrain its $E_p$, and truncation effects related to the instrument passband \citep{Lloyd00b}. The case of \swift\ BAT is an example of instrument affected by truncation biases: In spite of its higher fluence sensitivity than \batse, $E_p$ can be accurately determined only if it has a value within or very close to its energy passband (15--150 keV). However, we find some results about the validity of the $E_{p,i}$ -- $E_{\rm iso}$ relation untenable, such as the conclusion by \citet{Collazzi12}, who find that a significant fraction of the same GRBs with known redshift which have been used to derive the Amati relation do not satisfy their test. Clearly, their test condition (their so-called "Amati limit") in the (Fluence, $E_p$) plane is too restrictive. The same authors state that, while the Amati relation is the result of selection effects, the Ghirlanda relation is valid. Also, this statement is problematic, given that this relation is based on the same $E_p$ and Fluence measurements, being the only correction performed with, as discussed above, the replacement of $E_{\rm iso}$ with $E_{\gamma}$. In addition, the "Ghirlanda limit" is derived using rather loose assumptions about the beaming angle. The opposite conclusion was recently reached by \citet{Yonetoku10} via analyzing in detail all possible data truncation and detector sensitivity effects, and by \citet{Nava12}, who analyzed a complete sample, for redshift determination, of bright \swift\ GRBs (1~s peak photon flux $P \ge 2.6$~photons~s$^{-1}$~cm$^{-2}$ in the 15--150 keV BAT band). From this long standing debate, it is apparent the need to explore other approaches for testing the origin of the spectrum--energy correlations, i.e., whether they are the result of instrumental selection effects or are related to the fundamental physics of the GRB phenomenon. Given the significant evolution of the GRB spectra, studies of time-resolved spectra are crucial not only to test the $E_{p,i}$ versus $L_{\rm iso}$ relation, but also to delve deeper into the issue of the radiative mechanisms at work during the prompt emission. Motivated by both these needs, we performed a systematic study of the broad band (2--2000~keV) time--resolved prompt emission spectra of a sample of GRBs detected with both Wide Field Cameras (WFCs) aboard the \sax\ satellite and the \batse\ experiment aboard the {\em CGRO}. In this paper we will concentrate on the test of the $E_{p,i}$ versus $L_{\rm iso}$ relation. The WFCs were among the few instruments that detected GRB prompt emission down to 2 keV. Thus we can obtain time-resolved spectra in an energy band still not well explored: the 2--2000 keV band. A paper devoted to testing physical emission models of GRBs using the same time-resolved spectra will be the subject of a forthcoming paper (Frontera et al. 2012, in preparation).
|
By joining together the WFC and \batse\ spectral data of the four strongest ($>15\times 10^{-6}$~erg~cm$^{-2}$) GRBs simultaneously observed with both instruments, it was possible to perform a fine time-resolved spectral analysis in the broad energy band 2 keV to 2 MeV, a passband still scarcely explored as a whole. We do not expect significant systematic errors from this joint analysis. The response functions of both instruments are well known. Indeed, in the fits the cross-calibration factor was found to be always 1, in spite of being left to vary between 0.8 and 1.2 (see Section~\ref{analysis}). This is not the first time that a joint WFC/\batse\ spectral analysis has been performed. Results of similar analyses have been reported in the past by the \batse\ team \citep{Briggs00,Kippen03,Kippen04}. In addition, the \batse--deconvolved spectra of bright GRBs were cross-checked with those obtained with \sax\ GRBM \citep{Frontera09}; these in turn were cross-calibrated with WFC, with many published results \citep[e.g.][]{Frontera98a,Frontera00a}. For each of the strongest GRBs, we obtained a significant number of time-resolved spectra with constrained $E_p$: 8 spectra for GRB\,970111, 6 spectra for GRB\,980329, 19 spectra for GRB\,990123, and 7 spectra for GRB\,990510, with a total number of 40 analyzed spectra. With these spectra, we investigated the dependence of the time-resolved peak energy $E_p$ on the corresponding 2--2000 keV flux using the empirical BF as an input model. We find a significant power-law correlation between the derived peak energy $E_p$ and the flux within each GRB in our sample. The power-law best-fit parameters, evaluated with two different methods (the least-squares method, and the likelihood method with the addition of an external variance as a free parameter) give (see Table~\ref{t:ep-vs-flux}) similar index values in the cases of GRBs 970111 ($0.68\pm 0.06$ versus $0.65_{-0.14}^{+0.16}$) and 980329 ($0.16\pm 0.04$ versus $0.15_{-0.07}^{+0.08}$), and different values in the case of the GRBs 990123 ($0.53\pm 0.05$ versus $0.46_{-0.09}^{+0.09}$) and 990510 ($0.81\pm 0.15$ versus $0.56_{-0.23}^{+0.25}$), even if, in the latter case, these values are statistically almost consistent with each other. However, the likelihood method has the advantage of giving us information about the non-Poissonian spread of the data points around the best-fit curve. It shows that such a spread, even if small, affects the correlation, with the highest value ($\sigma_{\rm ext} = 0.09$~dex) in the case of GRB\,990510 and the minimum one ($\sigma_{\rm ext} = 0.00_{-0.00}^{+0.04}$~dex) in the case of GRB\,980329. Actually, in this case, also due to the low statistics of the data points, the correlation significance is low and, as shown by the reduced $\chi^2$ value (see Table~\ref{t:ep-vs-flux}), is not sensitive to an external spread. No clear correlation is found between the low-energy photon index $\alpha$ and flux and between $\alpha$ and $E_p$, apart from in one case (GRB\,980329). This means that the most relevant parameter that gives rise to the $E_p$--flux correlation is the break parameter $E_0$ in the BF. An equivalent important result is that the power-law correlation found between $E_p$ and flux is confirmed when we correlate, for two GRBs with known redshift, the intrinsic peak energy $E_{p,i}$ with the corresponding isotropic luminosity. Notice that the power-law index of the correlation changes from one GRB to the next when we use the least-squares method ($0.53\pm 0.05$ for GRB\,990123, and $0.81\pm 0.15$ for GRB\, 990510), but does not when the likelihood method is used ($0.46 \pm 0.09$ for GRB\,990123, $0.56_{-0.23}^{+0.25}$ for GRB\, 990510). This fact clearly means that the power-law index is affected by the external spread. If this spread is due to an unknown physical parameter, this result is an important hint for physical models of the prompt emission process: The intrinsic $E_{p,i}$ derived from the assumed emission model should be related to at least two physical parameters of the model. When we join together all the available data on GRBs with known redshift we find that the $E_{p,i}$--$L_{\rm iso}$ correlation becomes even more robust, with a probability of $1.57\times 10^{-13}$ that the correlation averaged over all data is due to chance, and a power-law index that is almost independent of the used best fit method ($0.66\pm 0.03$ in the case of the least-square method versus $0.63_{-0.07}^{+0.06}$ in the case of the likelihood method). It is also interesting to note that the non-Poissonian spread found for the GRB averaged correlation ($\sigma_{\rm ext} = 0.06_{-0.05}^{+0.06}$) is three times lower than the spread found in the Amati relation \citep{Amati09}, which is well known to be based on time-averaged spectra. This is a strong hint that part of the spread of the Amati relation is related with the fact that the $E_{p}$ values determined from time-averaged spectra are biased because of the spectral evolution of the GRB prompt emission. If we compare our GRB-averaged correlation result with that obtained by \citet{Ghirlanda10} from the time-resolved spectra of \fermi\ GRBs, we find that our results are consistent with those within a 2$\sigma$ belt, even if the slope obtained by these authors ($0.36 \pm 0.05$) is lower than that found by us ($0.63_{-0.07}^{+0.06}$; see Table~\ref{t:epi-vs-L} and Figure~\ref{f:epi-all-vs-L}), likely due to the sample variance. Indeed, our found slope is similar to that ($0.621 \pm 0.003$) reported by \citet{Lu12}, who performed the time-resolved spectral analysis of a sample of 15 \fermi\ GRBs with known redshift. Within a $2 \sigma$ spread, our results are also consistent with the time-averaged $E_{p,i}$ versus $L_{\rm iso}$ correlation, obtained by \citet{Ghirlanda10} using 95 pre-\fermi\ plus 10 \fermi\ GRBs (see the right panel of Figure~\ref{f:ep-vs-L-comp}). We cannot exclude that selection effects can influence our results, as with the fact that the analyzed GRBs are detected by both instruments, that they are sufficiently bright to allow time resolved analysis, and that they have known redshift. However, it seems difficult that this unavoidable selection can introduce a correlation between $E_p$ and $L_{\rm iso}$ within single GRBs. In addition, a similar correlation has been found by the other mentioned authors \citep{Ghirlanda10, Lu12} with other GRBs and with other instruments. To conclude our results strongly support, at least in the range of luminosities explored with our data, the reality of the Amati \citep{Amati02} and the Yonetoku \citep{Yonetoku04} relations, both derived using time averaged spectra. Also, our results give strong constraints on the physical models. In a forthocoming paper (Frontera et al. 2012, in preparation), with the same data we are testing different physical models, among them the recently developed {\sc grbcomp} model, which is devoted to the spectral formation of a GRB \citep{Titarchuk12}. In this model a physical interpretation of the Amati relation is also given.
| 12
| 6
|
1206.5626
|
1206
|
1206.2837_arXiv.txt
|
High-angular resolution observations of dense molecular cores show that these cores can be clumpier at smaller scales, and that some of these clumps can also be unbound or transient. The use of chemical models of the evolution of the molecular gas provides a way to probe the physical properties of the clouds. We study the properties of the clump and inter-clump medium in the starless CS core in LDN 673 by carrying out a molecular line survey with the IRAM 30-m telescope toward two clumps and two inter-clump positions. We also observed the 1.2-mm continuum with the MAMBO-II bolometer at IRAM. The dust continuum map shows four condensations, three of them centrally peaked, coinciding with previously identified sub-millimetre sources. We confirm that the denser clump of the region, $n\sim3.6 \times10^5$\cmt, is also the more chemically evolved, and it could still undergo further fragmentation. The inter-clump medium positions are denser than previously expected, likely $n\sim1\times10^3$--1$\times10^4$\cmt\ due to contamination, and are chemically young, similar to the gas in the lower density clump position. We argue that the density contrast between these positions and their general young chemical age would support the existence of transient clumps in the lower density material of the core. We were also able to find reasonable fits of the observationally derived chemical abundances to models of the chemistry of transient clumps.
|
\label{intro} It has been long known that molecular clouds are highly structured \citep[e.\, g.,][]{BlitzStark86} and their structure is greatly affected by the motions induced by supersonic turbulence \citep[e.\ g.,][]{Scalo98}, self-gravity of the gas and magnetic fields inside the clouds \citep{McKeeOstriker07,Kainulainen09}. All these processes control the formation and evolution of the density enhancements, of different scale sizes and densities, such as the cores and clumps that will finally give birth to stars. But there are still great uncertainties to identify the connection between starless cores and protostars \citep{Johnstone00,SmithClark08}. Starless cores are not all the same, despite their overall similarity in structure \citep{KetoCaselli08}, and differences in total mass, density and temperature might account for the differences in dynamical properties, structure, and future evolution of starless cores \citep{KetoField05,KetoCaselli08}. Higher-angular resolution observations are also finding that dense cores in molecular clouds that appeared homogeneous in single-dish observations, are clumpier at smaller scales \citep{Peng98,Morata03}, showing structures as small as 0.02 pc, and masses as low as 0.01\mo. Many of these smaller clumps are unbound and/or showing evidence of being transient \citep{Peng98,Morata05}, and will never be able to form low-mass stars or even brown dwarfs. The mix of bound and unbound structures in dense cores is also found in several regions, such as the Pipe Nebula, where recent molecular line and continuum observations \citep{Lombardi06,Muench07,Rathborne08,Frau10} found numerous cores more than 100, most of which appear to be pressure confined, and gravitationally unbound \citep{Lada08}. Observations of the emission of molecular lines at millimetre and sub-millimetre wavelengths in cloud cores combined with the modelling of the chemistry of the gas provides a way of obtaining information on the physical structure and the chemical and physical evolutionary stages of the cores. We proposed a time-dependent chemical model that also explored the consequences of the presence of unresolved and transient structures in the gas that would form and disperse in a timescale of $\sim1$--2 Myrs \citep{Taylor96}, in order to explain the systematic differences between CS and \nht\ lines \citep{Pastor91,Morata97}. Simulations of the evolution of these cores \citep{Garrod05,Garrod06} find that clouds that are ensembles of such transients have a clearly different chemistry from a `traditional' static cloud. The gas chemistry appears to be ``young'' at all times, and the re-cycling of the material frozen out onto dust grains produces a general molecular enrichment of the clouds, even after re-expansion of the transient structures. The background gas in which these inhomogeneities are embedded would be fairly diffuse, but chemically enriched. These chemical enhancements might also account for the variety of chemistries observed in diffuse clouds. We carried out interferometric high-angular resolution observations of several molecules (CS, \hcop\, and \ndh) \citep[][hereafter MGE03]{Morata03}, which we later combined with single-dish intermediate-angular resolution maps \citep[][from now on MGE05]{Morata05} towards the starless CS core in LDN~673 ($d=300$~pc), in order to test the predictions of the chemical models. The combined single-dish and interferometer maps showed emission of both background and clumped gas, with a clear segregation of clump properties between the northern and southern halves of our observed region, and allowed us to identify 15 resolved clumps in our data cube. The derived clump masses are well below the virial mass, which would point to their being transient, except for the more massive one, which might have a mass, $\sim 1$\mo, closer to the virial mass. The starless core appears to be constituted by a heterogeneous medium of condensations, of various densities and at different stages of chemical evolution, in agreement with theoretical studies that postulate the existence of transient clumps or the transient nature of dense cores generated by dynamical flows within molecular clouds \citep[see e.\, g.,][]{FH02, Vazquez05, VanLoo08}. Recently, \citet{Whyatt10} also found evidence of a heterogeneous medium in scales of less than 0.1 pc near HH objects, as traced by strong \hcop\ (3-2) emission. These clumps would have gas volume densities $\ga3\times10^4$\cmt. \begin{table} \caption{Lines observed with the 30-m IRAM telescope} \label{tablines} \centering \begin{tabular}{lccc} \hline\noalign{\smallskip} Transition & Frequency & HPBW & $B_\mathrm{eff}$ \\ & (GHz) & (arcsec) &\\ \noalign{\smallskip}\hline\noalign{\smallskip} CCH 1--0 \fj{3}{2}{1}{2} $F$=2--1 & \phn87.316925 & 28 & 0.78 \\ CCH 2--1 \fj{5}{2}{3}{2} $F$=3--2 & 174.663222 & 14 & 0.64 \\ CCH 3--2 \fj{7}{2}{5}{2} $F$=4--3 & 262.004260 & \phn9 & 0.46 \\ \noalign{\smallskip} CN 1--0 \jf{3}{2}{5}{2}--\jf{1}{2}{3}{2} & 113.490982 & 22 & 0.74 \\ CN 2--1 \jf{5}{2}{7}{2}--\jf{3}{2}{5}{2} & 226.874764 & 11 & 0.54 \\ \noalign{\smallskip} CS 5--4 & 244.935606 & 10 & 0.50 \\ \noalign{\smallskip} \cthd\ \jj{2}{1,2}{1}{0,1} & \phn85.338906 & 29 & 0.78 \\ \cthd\ \jj{4}{1,4}{3}{0,3} & 150.851899 & 16 & 0.68 \\ \noalign{\smallskip} HCN 3--2 & 265.886432 & \phn9 & 0.45 \\ \noalign{\smallskip} \hdco\ \jj{2}{1,2}{1}{1,1} & 140.839515 & 18 & 0.70 \\ \hdco\ \jj{3}{1,3}{2}{1,2} & 211.211448 & 12 & 0.57 \\ \hdco\ \jj{3}{1,2}{2}{1,1} & 225.697773 & 11 & 0.54 \\ \noalign{\smallskip} \htcop\ 1--0 & \phn86.754330 & 28 & 0.78 \\ \htcop\ 2--1 & 173.506782 & 14 & 0.64 \\ \htcop\ 3--2 & 260.255480 & 10 & 0.46 \\ \noalign{\smallskip} NO \jf{3}{2}{5}{2}--\jf{1}{2}{3}{2} $\Pi^+$ & 150.176459 & 16 & 0.68 \\ NO \jf{5}{2}{7}{2}--\jf{3}{2}{5}{2} $\Pi^+$ & 250.436845 & 10 & 0.48 \\ \noalign{\smallskip} SO \jj{3}{2}{2}{1} & \phn99.299905 & 25 & 0.76 \\ SO \jj{4}{5}{3}{4} & 206.176062 & 12 & 0.58 \\ SO \jj{6}{5}{5}{4} & 219.949433 & 11 & 0.55 \\ SO \jj{7}{6}{6}{5} & 261.843756 & 10 & 0.46 \\ \noalign{\smallskip} SO$_2$ \jj{3}{1,3}{2}{0,2} & 104.029410 & 24 & 0.76 \\ SO$_2$ \jj{7}{1,7}{6}{0,6} & 165.225436 & 15 & 0.66 \\ \noalign{\smallskip}\hline \end{tabular} \end{table} \begin{table} \caption{Coordinates of the positions selected for the spectral line observation.} \label{tabpositions} \begin{tabular}{llll} \hline\noalign{\smallskip} Position & Counterpart & R.A.(J2000) & Dec. (J2000)\\ \noalign{\smallskip}\hline\noalign{\smallskip} CL1 & \ndh\ peak & 19:20:51.747 & 11:13:49.50 \\ CL6 & CS peak & 19:20:50.003 & 11:14:53.00 \\ ICLN & Inter-clump N & 19:20:51.701 & 11:15:30.00 \\ ICLS & Inter-clump S & 19:20:54.501 & 11:14:10.00 \\ \noalign{\smallskip}\hline \end{tabular} \end{table} In order to study the properties of the clump and inter-clump gas in the starless CS core in LDN 673, we selected two positions associated with identified clumps (CL1 and CL6) and two positions where the inter-clump gas would be dominant (where we did not detect any clump). A multitransitional survey of several early- and late- type molecules in these positions allows us to sample the chemical composition of the gas and compare it to the predicted different chemistry of the pre- and post-clump gas in the models. We additionally observed the dust continuum emission in LDN~673. The structure of this paper is as follows: in Sect.~\ref{observations}, we describe the IRAM 30-m spectral line and continuum observations. In Sect.~\ref{results}, we describe the characteristics of the detected spectra and of the dust continuum emission. The analysis of the observational results and the determination of the physical parameters of the gas and dust are shown in Sect.~\ref{analysis}. Finally, Sect.~\ref{discussion} contains the discussion of the results of our analysis and how they can be related to the previous observations, the chemistry of the clouds and the structure of the core.
|
We studied the properties of the clump and inter-clump gas in the CS starless core of LDN~673 using the 30-m IRAM telescope through the emission of several spectral lines in the 3-, 2-, and 1-mm bands, in four positions associated with already detected, but physically and chemically different, clump gas (positions CL1 and CL6), or with inter-clump gas (positions ICLN and ICLS). We detected 19 spectral transitions of 10 molecular species at least in one of the four positions. We complemented the spectral observations with the mapping of the 1.2-mm dust continuum emission in LDN~673, which allowed us to obtain a more reliable estimate of the volume density of the gas and of the abundances of the detected molecules. The main results of our study are: \begin{enumerate} \item The dust continuum observations revealed four emission condensations in the region, roughly coinciding with SMM sources found by \citet{Visser02}, three of them with a round shape and centrally peaked emission: CL1/SMM5 is the most intense and coincides with the \ndh\ clump found by MGE03; SMM3, located $\sim50''$ to the West, was not previously found in the molecular observations; and SMM8, located $\sim2\farcm5$ SE of the centre of the map, and coinciding with an emission enhancement in the \cts\ and \htcop\ maps of MGE05. Finally, there is a more diffuse emission strip extended E--W $\sim2'$ north of CL1/SMM5. The masses of these condensations range from 1.6 to 5.1~\mo\ and sizes 0.05--0.07~pc. The northern condensation is sensibly more diffuse. \item We made a radiative transfer analysis using the RADEX code to determine the molecular column densities, volume density, and kinetical temperature at each position. The best fits to our observations found $T_\mathrm{K}\simeq10$~K for the CL1 and CL6, and $\simeq20$~K for the ICLN and ICLS positions; and volume densities ranging from $\sim3.6\times10^5$\cmt\ at CL1 to $\sim 1.7\times10^4$\cmt\ at ICLS. \item CL1 presents the largest column densities for most of the molecules (\htcop, \cthd, CCH, CN, and NO), but the smaller column densities for \hdco\ and SO. The estimated molecular abundances of most of the molecules, are smallest at CL1, while the abundances at CL6 and ICLN tend to be very similar. \item The comparison of the dust continuum emission to previous interferometer and single-dish observations shows very little or no correlation between the CS (2--1) and the 1.2-mm emission, with most of the clumps found in the CS emission not detected in the continuum observations. On the other hand, the dust continuum and \ndh\ emissions are much more similar. \item We found that the condensations detected in the dust continuum map may be subject to more thermal fragmentation, which might be already happening in CL1/SMM5. However, most of the clumps detected in CS, but undetected in the dust continuum observations, are unlikely to have been formed by thermal fragmentation, and would need other mechanisms to explain their formation. These clumps were proposed to be transient by MGE05. \item The chemistry of CL1 appears to be much more evolved than for the other positions, with signs of depletion of several molecular species (CS, \hdco, SO, \htcop, and partially NO), but also relatively unexpected high abundances of CCH, \cthd, and CN. The density contrast between CL6 and the two inter-clump positions, which are denser that initially expected, is relatively low. The gas at the CL6 and inter-clump positions seems to be generally chemically ``young''. \end{enumerate} In summary, the central condensation, CL1/SMM5, is probably approaching the 'peak-time' state of the models of \citet{Garrod05,Garrod06} and it is also the place with a larger probability of future undergoing of star formation (see MGE05). The SMM3 condensation is probably in a very similar state, from its shape and gas density, but we cannot firmly determine it until further observations provide us with more information about its chemistry. The low density contrast between CL6 and the inter-clump positions and their similar young chemical age seems to support the idea of the presence of lower density transient clumps in the core as proposed by \citet{Garrod06}.
| 12
| 6
|
1206.2837
|
1206
|
1206.2006_arXiv.txt
|
\\ A statistical study is carried out on the photospheric magnetic nonpotentiality in solar active regions and its relationship with associated flares. We select 2173 photospheric vector magnetograms from 1106 active regions observed by the Solar Magnetic Field Telescope at Huairou Solar Observing Station, National Astronomical Observatories of China, in the period of 1988--2008, which covers most of the 22nd and 23rd solar cycles. We have computed the mean planar magnetic shear angle ($\overline{\Delta\phi}$), mean shear angle of the vector magnetic field ($\overline{\Delta\psi}$), mean absolute vertical current density ($\overline{|J_{z}|}$), mean absolute current helicity density ($\overline{|h_{\rm c}|}$), absolute twist parameter ($|\alpha_{\rm av}|$), mean free magnetic energy density ($\overline{\rho_{\rm free}}$), effective distance of the longitudinal magnetic field ($d_{\rm E}$), and modified effective distance ($d_{\rm Em}$) of each photospheric vector magnetogram. Parameters $\overline{|h_{\rm c}|}$, $\overline{\rho_{\rm free}}$, and $d_{\rm Em}$ show higher correlation with the evolution of the solar cycle. The Pearson linear correlation coefficients between these three parameters and the yearly mean sunspot number are all larger than 0.59. Parameters $\overline{\Delta\phi}$, $\overline{\Delta\psi}$, $\overline{|J_{z}|}$, $|\alpha_{\rm av}|$, and $d_{\rm E}$ show only weak correlations with the solar cycle, though the nonpotentiality and the complexity of active regions are greater in the activity maximum periods than in the minimum periods. All of the eight parameters show positive correlations with the flare productivity of active regions, and the combination of different nonpotentiality parameters may be effective in predicting the flaring probability of active regions.
|
\label{S-introduction} It is generally accepted that the magnetic field dominates the evolution of the Sun, and the magnetic nonpotentiality may be responsible for massive energy released in solar explosive activities such as flares and coronal mass ejections (CMEs). Even though the physical parameters that characterize the nonpotentiality of the magnetic fields do not directly trigger solar eruption, active regions (ARs) with strong nonpotentiality and considerable complexity are more likely to explode than simpler ones because the former have sufficient free energy to release. Studying the evolution of magnetic nonpotentiality and its relationship with solar eruption from a statistical point of view is useful in predicting eruptive events and in understanding the long-term evolution of solar activity. The characteristics of nonpotential magnetic fields associated with flares in ARs have been studied for a couple of decades. Several nonpotential parameters of flare-productive ARs are investigated by many authors in detail. \inlinecite{Gary87} analyzed nonpotential features based on the observed vector magnetic field of NOAA AR 2684, and tried to answer the questions such as the relation between the nonpotential features and the flares. \inlinecite{WangJX96} examined in detail the relationship between flare occurrence and nonpotentiality development in a flare-productive region. \inlinecite{Moon00} analyzed the spatial distribution and temporal evolution of several nonpotential parameters in NOAA AR 5747. \inlinecite{Deng01} studied the daily evolution of magnetic nonpotentiality focusing on NOAA AR 9077 that produced a giant flare from 11 to 15 July 2000. \inlinecite{Leka03} investigated the magnitudes and temporal variations of several photospheric magnetic parameters, such as horizontal magnetic gradient, vertical current, current helicity, twist parameter $\alpha$, magnetic shear angles, and excess magnetic energy in three ARs, to identify properties that are important to the solar flare production. \inlinecite{Dun07} calculated the magnetic shear angle, vertical current, and current helicity of NOAA AR 10486 and analyzed their temporal evolutions and spatial relations to the flares. \citeauthor{Falconer02} (\citeyear{Falconer02}, \citeyear{Falconer06}) also discussed the relationship between several nonpotentiality-related parameters and CME productivity of ARs. Some authors studied the physical characteristics of photospheric magnetic fields in ARs and their relationship to flares in a statistical sense. \inlinecite{Cui06} preferred the maximum horizontal gradient, length of the neutral lines, and number of singular points; \inlinecite{Jing06} used the mean value of magnetic gradients at strong-gradient magnetic neutral lines, length of strong-gradient magnetic neutral lines, and total magnetic energy; \inlinecite{Cui07} also tried the length of neutral lines with strong gradient and/or strong shear; \inlinecite{Cui08} studied the total unsigned current and total unsigned current helicity. \citeauthor{Guo06} (\citeyear{Guo06}, \citeyear{Guo07a}, \citeyear{Guo10}) utilized the effective distance that could roughly quantify the magnetic complexity of ARs, to study the evolution of magnetic complexity in ARs, its relationship with erupting activity of ARs, and its evolution with the solar cycles, respectively. A close relationship between magnetic nonpotentiality of ARs and solar eruptions is observationally studied in numerous articles. \inlinecite{Rust94} reviewed the progress of pre-flare state studies, including nonpotential characteristics, and pointed out the importance of the more quantitative results from the vector magnetographs. Lacking of long-term steady vector magnetic field observations, the line-of-sight magnetograms are mostly used as means of case analyses and statistical researches. It is necessary to study the long-term evolution of appropriate parameters derived from vector magnetic fields and find whether there is a correlation between these parameters and flares or other eruption events. So far ground-based telescopes, whose maintenance for long-term observations is relatively easy, still have the advantages in data accumulation. Since 1987, the Solar Magnetic Field Telescope (SMFT) \cite{Ai86}, installed at Huairou Solar Observing Station (HSOS), National Astronomical Observatories of China (NAOC), has reliably worked for more than twenty years that almost covers two complete solar cycles. Based on the long-term observations of SMFT, a statistical study on photospheric magnetic nonpotential parameters in ARs is carried out, and the relationship between these parameters and solar flares is also presented. In Section \ref{S-observation}, observations by SMFT and data reductions are introduced. Measures of nonpotentiality and complexity are shown in Section \ref{S-param}. In Section \ref{S-analysis}, statistical analysis and results on magnetic nonpotentiality associated with flares are presented. Conclusions are given in Section \ref{S-conclusion}.
|
\label{S-conclusion} By calculating eight parameters ($\overline{\Delta\phi}$, $\overline{\Delta\psi}$, $\overline{|J_{z}|}$, $\overline{|h_{\rm c}|}$, $|\alpha_{\rm av}|$, $\overline{\rho_{\rm free}}$, $d_{\rm E}$, and $d_{\rm Em}$) of nonpotentiality and complexity for 2173 photospheric vector magnetograms in 1106 ARs associated with flares, we found the main results as follows: \begin{description} \item (1) On average, two mean magnetic shear angles $\overline{\Delta\phi}$ and $\overline{\Delta\psi}$, mean absolute vertical current density $\overline{|J_{z}|}$, absolute twist factor $|\alpha_{\rm av}|$, and effective distance $d_{\rm E}$ in ARs do not change significantly with the global solar activity level. However, it is more likely that these parameters show higher values in the solar maximum than in the solar minimum. \item (2) The mean absolute current helicity density $\overline{|h_{\rm c}|}$, mean free magnetic energy density $\overline{\rho_{\rm free}}$, and modified effective distance $d_{\rm Em}$ show high positive correlation with the mean sunspot number, and these parameters also have relatively close relationship with each other. The Pearson linear correlation coefficients of the above three with the yearly mean sunspot numbers are larger than 0.59. They can be used to characterize the solar activity level as well as the traditional sunspot number. \item (3) The nonpotentiality and complexity parameters between flare-productive ARs and flare-quiet ones computed in this work are useful to understand the evolution of flare-productive ARs and the relationship with the magnetic activity levels of the cycles. These nonpotentiality and complexity parameters may be synthetically applied as indicators to predict solar flares with some weight. \item (4) Due to the loss of the information of magnetic field strength in the parameter of effective distance $d_{\rm E}$, the modified effective distance $d_{\rm Em}$ (including the strength of the magnetic field) turns out to be much better in indicating the magnetic activities of ARs. \end{description} The accumulated nonpotential energy in the magnetic field of solar ARs provides the energy of solar flares, while the trigger of flares is also related to the magnetic nonpotentiality. The parameters discussed above will provide the basic information of nonpotentiality of solar active regions in different aspects. The synoptic analysis of different nonpotential parameters may supply an objective basis on the relationship between the nonpotential field and solar flares in order to predict solar flares. Although long-term accumulation of the data from ground-based instruments is indispensable, the vector magnetograms taken by the {\it Hinode} satellite, {\it Solar Dynamics Observatory} (SDO), and other space-borne solar telescopes without being affected from atmospheric seeing have been giving more accurate data with better spatial resolution for further comparison studies. \begin{acks} The authors wish to thank the referee and editor for important comments and suggestions. One of the authors X.Yang is grateful to Prof. X.J.Mao and Dr. J.Jiang for reading through the manuscript and giving beneficial advices. This work is supported by the National Natural Science Foundation of China (10733020, 10921303, 11003025, 11103037, 11103038, 60940030, 11173033, and 41174153), the Young Researcher Grant of National Astronomical Observatories of Chinese Academy of Sciences (CAS), the Knowledge Innovation Program of CAS (KJCX2-EW-T07), and the Key Laboratory of Solar Activity of CAS. The authors are grateful to Mrs. G.P.Wang and the HSOS staff for producing the nice data, and also the GOES team for the helpful records. \end{acks}
| 12
| 6
|
1206.2006
|
1206
|
1206.1005_arXiv.txt
|
Solar flare X-ray emission results from rapidly increasing temperatures and emission measures in flaring active region loops. To date, observations from the X-Ray Sensor (XRS) onboard the {\it Geostationary Operational Environmental Satellite (GOES)} have been used to derive these properties, but have been limited by a number of factors, including the lack of a consistent background subtraction method capable of being automatically applied to large numbers of flares. In this paper, we describe an automated temperature and emission measure-based background subtraction method (TEBBS), which builds on the methods of \cite{born90}. Our algorithm ensures that the derived temperature is always greater than the instrumental limit and the pre-flare background temperature, and that the temperature and emission measure are increasing during the flare rise phase. Additionally, TEBBS utilizes the improved estimates of \goes temperatures and emission measures from \cite{whit05}. TEBBS was successfully applied to over 50,000 solar flares occurring over nearly three solar cycles (1980-2007), and used to create an extensive catalog of the solar flare thermal properties. We confirm that the peak emission measure and total radiative losses scale with background subtracted \goes X-ray flux as power-laws, while the peak temperature scales logarithmically. As expected, the peak emission measure shows an increasing trend with peak temperature, although the total radiative losses do not. While these results are comparable to previous studies, we find that flares of a given \goes class have lower peak temperatures and higher peak emission measures than previously reported. The resulting TEBBS database of thermal flare plasma properties is publicly available on Solar Monitor (www.solarmonitor.org/TEBBS/) and will be available on Heliophysics Integrated Observatory (www.helio-vo.eu).
|
\label{sec:intro} Solar flares are among the most powerful events in the solar system, releasing up to $10^{33}$ ergs in a few hours or even minutes. They are believed to be powered by magnetic reconnection, a process whereby energy stored in coronal magnetic fields is suddenly released. According to the CSHKP flare model \citep{carm64,stur66,hira74,kopp76}, electrons accelerated by magnetic reconnection spiral down the magnetic loops and strike the chromosphere causing the emission of hard X-rays (HXR). As a consequence, the chromospheric material is also heated and expands back up into the loops which causes the observed increase in temperature and emission measure \citep[e.g.,][]{flet11}. To date, the study of solar flares has been predominantly focused on single events or small samples of events. While such studies have furthered our understanding of the physics of these particular flares, they are fundamentally limited since they cannot, with any certainty, explain the global behavior of solar flares. In contrast, only the study of large-scale samples can give an insight as to whether findings of given studies are particular to individual events or characteristic of many. This can allow constraints to be placed on global flare properties and give a greater understanding to the fundamental processes which drive these explosive phenomena. That said, large-scale studies of solar flare properties have been few in number over the past decades. Such a study was performed by \cite{garc92} who used the Geostationary Operational Environmental Satellite (\goess) to examine 710 M- and X-class flares. They noted a sharp linear lower bound in the relationship between emission measure and \goes class. However, this paper is mainly focused on categorizing types of very high temperature flares and examined whether these flares approached or exceeded this emission measure lower bound. A definitive example of a large-scale study of the thermal properties of solar flares was conducted by \cite{feld96b}, who combined results from three previous studies \citep{phil95,feld95,feld96a} to investigate how temperature and emission measure vary with respect to \goes class for 868 flares, from A2 to X2. Their work used temperatures derived using the Bragg Crystal Spectrometer (BCS) onboard \emph{Yohkoh}. These temperature values were convolved with the corresponding \goes data to derive values of emission measure. They found a logarithmic relationship between \goes class and temperature, and a power-law relationship between \goes class and emission measure, with larger flares exhibiting higher temperatures and emission measures. However, temperature and emission measure were derived at the time of the peak 1--8~\AA~flux and so are likely to be less than their true maxima. Furthermore, BCS temperatures have been found to be higher than those measured by \goes \citep{feld96b}, and using these values to calculate \goes emission measure will give lower values than if \goes was used consistently. More recently, \cite{batt05} studied the correlation between temperature and \goes class for a sample of 85 flares, ranging from B1 to M6 class. Although the values reported gave a flatter dependence than \cite{feld96b}, the large scatter in the data led to a very large uncertainty making the two relations comparable. In contrast to \citet{feld96b}, \citet{batt05} accounted for solar background and extracted the flare temperature at the time of the HXR burst as measured by the {\it Ramaty High-Energy Solar Spectroscopic Imager} (\rhessis; \citealt{lin02}) rather than at the time of the soft X-ray (SXR) peak. However, any discrepancies expected to be caused by these differences were not discernible in view of the large uncertainties. In addition, \citet{casp10} used \rhessi to examine the temperature of 37 high temperature flares with \goes class. The relationship was qualitatively similar to those of \citet{feld96b} and \citet{batt05} however as the relationship was not fit no quantitative comparison can be made. Larger statistical samples were studied by \cite{chri08} and \cite{hann08}, who investigated the frequency distributions and energetics of 25,705 microflares (\goes class A--C) observed by \rhessi from 2002 to 2007. From those events for which an adequate background subtraction could be performed (6,740) a median temperature of $\sim$13~MK and emission measure of 3$\times$10$^{46}$~cm$^{-3}$ were found. \cite{hann08}, in particular, looked at the temperature derived from \rhessi observations as a function of (background subtracted) \goes class, and found similar trends to the works of \cite{feld96b} and \cite{batt05}. However, their analysis only included events of low C-class and below. While these studies have provided some insight into the global properties of solar flares, they each have their limitations. In particular they lack a commonly used method of isolating the flare signal from the solar and instrumental background contributions. Previous background subtraction methods have often been performed manually. Others, such as setting the background to the flare's initial flux values, or fitting polynomials between the flux values at the start and end of the flare, often exaggerate noise and do not preserve characteristic temperature and emission measure evolution. Therefore, the accurate separation of flare signal and background limits the number of events that can be analyzed. For example, \cite{batt05}, in accounting for solar background, were only able to compile a sample of 85 events. Although a larger dataset would not have reduced the range of scatter, it would have better revealed the variations in the density of points within the distribution. This would have allowed a fit to be more tightly constrained and thereby reduced the uncertainties. Conversely, \cite{feld96b}, with a sample of hundreds of flares, did not attempt to account for the solar background at all, which can bias smaller events as the background makes up a greater contribution to the overall flux. Few attempts have been made to develop automated background subtraction techniques for \goes observations which can be applied to large numbers of flares. \cite{born90} developed a method to determine whether a given background subtraction preserves characteristic temperature and emission measure evolution without checking manually, i.e., that temperature and emission measure both increase during the rise phase of flares. This behaviour has been seen in numerous observations (e.g., \citealt{flud95,batt09}) and numerical models (e.g., \citealt{fisc85,asch09}). This method used the polynomials of \citet{thom85} which relate temperature and emission measure to the ratio of the short and long \goes channels, $R=F_S/F_L$. However, \cite{whit05} have since improved on this by assuming more modern spectral models \citep[CHIANTI 4.2][]{land99,land02} and taking into account the differences between coronal and photospheric abundances, requiring the tests of \cite{born90} to be updated. In this paper, we study the thermal properties of solar flares using \goes observations over nearly three solar cycles. The flare signal within these observations has been isolated from the various solar, non-solar, and instrumental background contributions using a modified background subtraction method. This in turn has allowed more accurate automatic calculation of flare properties. In Section~\ref{sec:goes} of this paper, we discuss the \goes XRS, the \goes event list and how to derive plasma properties from \goes observations. In Section~\ref{sec:method} we describe previous background subtraction methods for \goes observations and outline how we have improved upon the work of \cite{born90}. In Section~\ref{sec:results} we use this method to improve upon previous statistical studies by deriving flare properties such as peak temperature and emission measure for flares in the \goes event list and examining the relationships between them. In Sections~\ref{sec:discussion} and \ref{sec:conclusions} we discuss the results and provide some conclusions.
|
\label{sec:conclusions} A method, TEBBS, has been presented for isolating the solar flare signal from \goes soft X-ray lightcurves by accounting for contributions from both solar and non-solar backgrounds. This allows the properties of the flaring plasma itself to be more accurately derived. It can be systematically applied to any number of flares, removing many of the inconsistencies that can be introduced when manually defining a background level. This makes it a particularly suitable method for conducting large-scale statistical studies of solar flares characteristics. TEBBS was found to produce fewer spurious artifacts in the derived temperature and emission measure profiles for both individual events (Figure~\ref{fig:threemethod}) and in large statistical samples (Figures~\ref{fig:plasma}~and~\ref{fig:t_vs_em}), compared to when either all or none of the pre-flare flux was removed. This led to more reliable relationships being derived between flare plasma properties (temperature, emission measure etc.), which can in turn place constraints on the `allowed' values of properties for a flare of a given GOES magnitude. TEBBS was successfully applied to 50,056 flares from B-class to X-class, making it the largest study of the thermal properties of solar flares to date. It was found that peak temperature scales logarithmically with peak long channel flux as described by Equation~\ref{eqn:t(f)fit}. Meanwhile, peak emission measure and total radiative losses scaled with peak long channel flux as power-laws given by Equations~\ref{eqn:em(f)fit} and \ref{eqn:lradfit}. Uncertainties were calculated for these derived relations unlike previous studies. The exception to this was \citet{batt05} who provided uncertainties for their slopes only. The uncertainties derived using TEBBS were nonetheless smaller than those of \citet{batt05} and include uncertainties on the intercepts as well as slopes. Furthermore, while these results are broadly in line with previous studies, it was found that flares of a given \goes class have lower temperatures and higher peak emission measures than previously reported. Peak emission measure and total radiative losses were also examined as a function of peak temperature. It was found that flares with high peak temperatures also have high peak emission measures (in agreement with \citealt{garc88} and \citealt{garc92}). However, the derived correlation was relatively weak. Similarly, it was also found that flares of a given peak temperature could exhibit a large range of radiative losses with no clearly defined trend. This lack of a clearly defined relationship between two derived properties could be attributed to the assumptions that go in to calculating them. Although both a constant density and a fixed coronal abundance were assumed in this study, both have been shown to vary during individual events \citep[e.g.][]{grah11}. A followup analysis of how changes in these variables might affect the derived properties, particularly in conjunction with hydrodynamic simulations, may lead to more reasonable correlations. This compilation of solar flare properties represents a valuable resource from which to conduct future large-scale statistical studies of flare plasma properties. For example, \cite{stoi08} derived analytical predictions of temperature and emission measure in response to electron beam and conduction driven heating and compared the results to \rhessi observations of 18 microflares. They found an order of magnitude discrepancy between conduction driven emission measures predicted by the Rosner-Tucker-Vaiana (RTV; \citealt{rtv78}) scaling laws and observation. This seemed to suggest that electron beam processes dominated. However, they noted that \rhessis's high temperature sensitivity ($\gtrsim$10~MK) mean that the observed temperatures may not have well represented the conduction value of the microflares, thus explaining the discrepancy. The fact that the \goes XRS has a sensitivity to lower temperatures than \rhessi makes the TEBBS database ideal for exploring this possibility. Since the RTV scaling laws and electron beam heating models are widely used to understand and model solar flares, it is important to examine disagreements between their predictions and observation. Another example of the use of RTV scaling laws in understanding flares is \cite{asch08}. They used these laws to derive theoretical ($EM \propto T^{4.3}$) and observed ($EM \propto T^{4.7}$) scaling laws between peak temperature and emission measure for solar and stellar flares. However, as part of their study, results from previous studies such as \cite{feld96b} and \cite{feld95} were included which did not account for background issues. TEBBS can therefore also be used to examine these scaling laws with greater statistical certainty and therefore provide more clarity on the discrepancies between theory and observation. As the scaling laws derived by \cite{asch08} apply to solar and stellar flares, conclusions drawn from TEBBS can be extended to stellar flares as well. TEBBS can be used to examine a wide range of flare characteristics, such as thermodynamic evolution and, in light of the work of \cite{stoi08}, even flare loop topologies. As TEBBS is also the largest database of thermal flare plasma properties to date, it will provide a valuable resource for future solar flare research.
| 12
| 6
|
1206.1005
|
1206
|
1206.1834_arXiv.txt
|
It is suggested in observations of supernova remnants that a number of large- and small-scale structures form at various points in the explosion. Multidimensional modeling of core-collapse supernovae has been undertaken since SN1987A, and both simulations and observations suggest/show that Rayleigh-Taylor instabilities during the explosion is a main driver for the formation of structure in the remnants. We present a case study of structure formation in 3D in a \msol{15} supernova for different parameters. We investigate the effect of moderate asymmetries and different resolutions of the formation and morphology of the RT unstable region, and take first steps at determining typical physical quantities (size, composition) of arising clumps. We find that in this progenitor the major RT unstable region develops at the He/OC interface for all cases considered. The RT instabilities result in clumps that are overdense by 1-2 orders of magnitude with respect to the ambient gas, have size scales on the level of a few \% of the remnant diameter, and are not diffused after the first $\sim30$ yrs of the remnant evolution, in the absence of a surrounding medium.
|
} Morphological, kinematic, and compositional structures are ubiquitous in the observations of supernovae (SNe) and supernova remnants (SNRs). These structures span scales from unipolar asymmetries across the whole remnant to sub-AU (astronomical unit) sized high density knots being shredded in the reverse shock of the Cassiopeia A (Cas A) SNR. The dense knots in SNRs are of particular interest from a nucleosynthetic and astrobiological point of view as vehicles for the chemical enrichment of star and planet-forming material in high mass star formation regions, as well as for interpreting observations of remnants. They carry nearly undiluted material from the metal-rich mantle of the former star and thus are good candidates for studying hydrodynamics and mixing processes during the explosion with both numerical and observational tools. Multiple physical processes drive structure formation. Although it had long been known that instabilities would grow in the shock launched in a supernova explosion (Chevalier 1976), most explosion studies focused on 1-dimensional models (primarily due to the high computational requirements of multi-dimensional simulations). But SN 1987A demonstrated the wide variety of observables affected by these instabilities: broad line widths in the infrared and gamma-ray lines of several elements \citep{Erickson88,Witteborn89}, low velocity hydrogen features in the spectrum 221 days after the explosion \citep{Hoflich88}, and indirect evidence from light curve models \citep{Woosley88,Shigeyama88,AFM89}. All these observables suggested deep mixing had occurred in the explosion \citep[see][for a review]{mix07}. While turbulence occurs during many stages of the collapse and explosion process of a massive star, the particular focus of this paper are the instabilities caused by the interaction of the SN shock wave with steep gradients in the profile of the exploding star. \citet{AFM89} reviewed a number of sites/events in an exploding star that can lead to deviations from spherical behavior, and point out that the formation of Rayleigh-Taylor (RT) fingers by shock passage as the most important. RT instabilities arise commonly in situations where a less dense fluid is accelerated into a denser fluid (e.g. when a less dense fluid is supporting a denser fluid against gravity), or more generally, where a fluid of higher entropy is accelerated into one with lower entropy. In the limit of impulsive acceleration this is referred to as a Richtmyer-Meshkov (RM) instability. Bubbles of the higher entropy fluid rise into the less entropic fluid, while columns or spikes of that penetrate into the higher entropy fluid. Shear flows at the interface between the two fluids are subject to Kelvin-Helmholtz (KH) instabilities. In the case of RM instabilities, both scenarios, a shock accelerated into an interface going from heavy to light and light to heavy fluids are unstable in the RM sense. The RM instability results in very similar looking features as the RT instability. It is likely that both instabilities are occurring during the explosion, and distinguishing between them may be somewhat subjective. For ease of reading we will refer to the whole class of instabilities henceforth as RT instabilities unless the distinction makes an important difference in the interpretation. In computer simulation of the shock propagation through the star, multiple sites have been found to become unstable and result in the growth of RT instabilities. Nearly all simulations to date \citep{FAM91,MFA91,HB92, HB91,Hachisu_ea91,Hachisu_ea92,NSS98, Kifon_ea03,Kifon_ea06,HFW03, HFR05,JWH09,JAW10,HJM10} find that strong instabilities grow. The RT instabilities most often arise at the He/metals interface for different progenitor models, which typically resulted in the mixing down of H and He, and the mixing out of at least C and O, and often higher- A elements like $^{24}$Mg, Si, and the Fe-group, though not always in sufficient quantities to explain observations on SN1987A. A higher degree of non-linearity in the RT instabilities can be achieved with an asymmetric shock front \citep{Hachisu_ea92}. \citet{FAM91} showed that, except for very coarse grids, the mode of the instability (i.e. the average spacing between RT fingers) is independent of resolution. \citet{Hachisu_ea92} and \citet{HB91} also showed that the amplitude of artificial seed perturbations (which are imposed to a) counter the damping of the highest modes due numerical and/or artificial viscosity, and b) to mimic fluctuations which are likely present in physical stars) does not influence the RT instabilities significantly, as long as there is a perturbation. Often more than one region becomes unstable in simulations, and the different instabilities then in many cases interact and merge. Extending their study to higher resolution, \citet{MFA91} discovered that the RT fingers first form at the H/He interface, but is then overrun by RT forming at the He/metals interface. Using a slightly different approach, \citet{HB91,HB92} modeled 1987A with a particle-based numerical scheme, and obtained similar results. Depending on the progenitor used, multiple sites became RT unstable, which in some case merged to just one instability. \citet{Muller_ea89} also emphasized the importance of using an accurate stellar density profile, since the polytropic profile in their earlier calculations showed no evidence for unstable regions. \citet{HB92} in a sense expanded on this conclusion by demonstrating the different RT morphologies achieved with different progenitor star profiles. These early results \citep[e.g.][]{MFA91, HB92,Hachisu_ea92}, though, suggested that such mixing as the shock moves through the star was insufficient to explain the mixing in SN 1987A. To enhance this transport, scientists revived research studying initial perturbations from convection in stellar progenitors\citep{BazanArnett98,Kane_ea00} and in the explosion by studying aspherical effects in the core-collapse engine\citep{Herant92,Herant_ea94}. In the core-collapse engine, these studies showed that turbulence above the proto-neutron star is important in producing an explosion. Although there are disagreements as to the nature of the instabilities (standing accretion shock vs. Rayleigh Taylor, etc.) this convection-enhanced engine is the current favored model in core collapse\citep{Herant_ea94,Burrows95, Mezz_ea98,FryerWarren02,Blondin03,Buras03,Blondin06,Burrows06,FY07}. This convective engine can produce highly asymmetric explosions. Such asymmetries will drive mixing as the shock moves out of the star. \citet{HFW03,HFR05} studied the effect of these explosion asymmetries on the mixing using 3-dimensional models. Their results showed that artificially imparted explosion asymmetries can dominate the mixing, producing broad line profiles like those in SN 1987A (where symmetric runs of the same explosion energy could not). This work found that $^{56}$Ni was mixed well into the hydrogen layer for the most asymmetric explosions and argued that the asymmetries could explain both the rapid rise in gamma-ray radiation as well as the redshift of the gamma-ray emission. Realizing the importance of perturbations set up by the shock revival mechanism, multi-dimensional explosion calculations are now being used for shock propagation calculations. \citet{Kifon_ea03} followed the explosion mechanism and the propagation of the blast wave simultaneously in 2D. RT instabilities, during the early convection that revived the shock, resulted in a slightly aspherical distribution of $^{56}$Ni. This distribution imprinted long-wavelength perturbations on the Si/O layer, and out of which RT instabilities grew as that interface became unstable. RT instabilities were also observed at the He/CO interface. They also found that the deeper RT instability at the Si/O interface resulted in the mixing out of some Ni. \citet{HJM10} use a 3D explosion calculation from Scheck (2007) to follow the shock propagation through a 15.5M$_\odot$ blue supergiant star in 2D and 3D under differing initial conditions. This is one of the first calculations to follow both the launch of the shock and the ensuing explosion in 3D. Some slight deformation from sphericity by the supernova engine seeds the later growing RT instabilities in their simulations (no artificial seed perturbations were implemented), with the sites of the largest deformation resulting in the largest RT plumes. RT fingers again formed at the He/CO interface, and also at the Si/O interface, and fragmented into clumps. \citet{JWH09} presented simulations for a small number of progenitors - 2 masses and 2 metallicities - in 2D, and \citet{JAW10} extend that study to 3D, though in the interest of saving computational resources all explosion models were initiated in 1D. Prominent RT instabilities develop again at the He/O interface, though instabilities at the Si/O are possibly suppressed due to the explosion mechanism used. The more massive progenitor in each case showed a wider region of instabilities, and \citet{JWH09} state that in the solar metallicity, 25M$_\odot$ case, RT instabilities extended down past the O shell and into the Si/S layer, resulting in increased mixing out of Fe-group material. They, again, find that the profile of the specific progenitor has a large influence on the extent and morphology of the RT region. The formation of RT and RM instabilities by shock interaction with interfaces thus is a robust feature in supernova explosion simulations. Some variance in the details and location of the RT and RM instabilities exists between different calculations, mostly due to the uncertainty in post-main sequence stellar structure and due to different explosion algorithms used. Previous calculations were generally more focused on the emergence of and mixing that these instabilities produce, as they were often compared to SN1987A. However, in order to be able to make meaningful comparisons to older remnants like Cas A, where the evolution is dominated by the interaction of the ejecta with surrounding stellar winds and/or ISM, it is necessary to extend these calculations to a much longer time after shock breakout. Among other things, this can shed more light on the further evolution of the clumps created by RT instabilities, and help firmly establish their relationship to features like the dense ejecta knots in Cas A. It is our aim to follow structures from their formation all the way to the young remnant phase, e.g. similar in age to Cas A. Identifying the location and timing of structure formation and modification, and comparing them to observations of young SNR will elucidate the proximate deposition of nucleosynthesis products in the interstellar medium of star and planet formation regions and the history of SNRs. In this paper we present the first step in this endeavor, 3 dimensional simulations of a \Msol{15} SN explosion evolved out to the homologous expansion phase. This will establish our methodology for simulations out to later times with circumstellar medium interactions and comparison of different progenitors and explosion asymmetries. We also propose a method of characterizing the sizes of overdense clumps that can be compared directly with observations. In section \ref{s:sims} we describe our simulations and the parameters that we explored. We offer an analysis useful for observational comparison for determining typical clump sizes in section \ref{s:methods}. A discussion of our results is presented in section \ref{s:results}, and some concluding remarks are presented in section \ref{s:conclusion}
| 12
| 6
|
1206.1834
|
|
1206
|
1206.5161_arXiv.txt
|
We clarify how magnetic reconnection can be derived from magnetohydrodynamics (MHD) equations in a way that is easily understandable to university students. The essential mechanism governing the time evolution of the magnetic field is diffusion dynamics. The magnetic field is represented by two components. It is clarified that the diffusion of a component causes a generation of another component that is initially zero and, accordingly, that the magnetic force lines are reconnected. For this reconnection to occur correctly, the initial magnetic field must be directed oppositely in the two regions, e.g., $y>0$ and $y<0$; must be concave (convex) for $y>0$ ($y<0$); and must be saturated for $y$ far from the $x$ axis, which would indicate the existence of the current sheet. It will be clear that our comprehension based on diffusion runs parallel to the common qualitative explanation about the magnetic reconnection.
|
A solar flare is the most violent explosion in the solar system. It occurs over the Sun's surface, resulting in the brightening of electromagnetic waves over a wide wavelength range and the release of about $10^{22}\mbox{--}10^{25}$ joule of energy within about $10^2\mbox{--}10^3$ s\cite{Yokoyama10}. How such a large amount of energy is released in such a short time has puzzled scientists for many years. One of the epoch-making approaches to this problem is to exchange the magnetic energy into kinetic energy and the heat of plasma by reconnecting magnetic field lines. This is referred to as {\em magnetic reconnection}. The time evolution of the magnetic field is determined by a magnetic induction equation that includes two contrary effects, ``freezing'' and diffusion of the magnetic field, which is explained in detail later. Diffusion of the magnetic field is caused by Joule dissipation. If the electric resistivity is low, the magnetic field is ``frozen'' into the plasma. However, if an anomalous resistivity is induced for some reason and the diffusion effect becomes dominant, the freezing of the magnetic field ceases and its structure can be altered. In other words, the diffusion causes structural change of the magnetic field. However, the time for the release of the magnetic energy is quite long. In the case of the solar flare, the diffusion constant $\eta$ is about $1 \ \mbox{m}^2/\mbox{s}$ and the typical scale of length for the energy release, which we shall denote by $L_0$, is $10^8 \ \mbox{m}$. Therefore, the time can be estimated as ${L_0}^2/\eta\sim 10^{16} \ \mbox{s} \sim 10^{8} \ \mbox{yr}$, which is not comparable with the observed release time\cite{Priest}. Hence, we can explain the heating of plasma and the release of a large amount of energy by considering only the time evolution of the magnetic field, but cannot explain the release time. For resolving this problem, Dungey proposed a model considering the electromagnetic field caused by the motion of the plasma and, thus, using magnetohydrodynamics (MHD) with negligible gas pressure. He showed that a current sheet forms between opposite directions of the magnetic field\cite{Dungey53,Priest}. Dungey was also the first to suggest that the configuration of the magnetic field can change. Subsequently, Sweet and Parker expanded on Dungey's model by considering the in/out flow of the plasma and using MHD equations; this model is referred to as {\em Sweet-Parker's model}\cite{Sweet58a,Sweet58b,Parker57,Priest}. According to their model, a plasma flowing into the current sheet with a velocity $v_{i}$ flows out along the sheet with the Alfv\'{e}n velocity $V_A$. The velocity $v_{i}$ satisfies the following relation. \begin{equation} v_i = {R_m}^{-1/2}V_A \ , \end{equation} where ${R_m}$ denotes the magnetic Reynolds number, which can be expressed as \begin{equation} R_m = \frac{L_0V_A}{\eta} \ . \end{equation} The magnetic energy moves with velocity $v_{i}$, and the time with which the energy is released over a length $L_0$ is given by \begin{equation} \frac{L_0}{v_i} = \frac{{L_0}^{3/2}}{\sqrt{\eta V_A}} \ . \end{equation} If we use $10^8 \ \mbox{m}$ and $1 \ \mbox{m}^2/\mbox{s}$ as values for $L_0$ and $\eta$, respectively, and estimate the Alfv\'{e}n velocity as $10^{8} \ \mbox{m}/\mbox{s}$, the energy release time becomes $10^8 \ \mbox{s}$. Therefore, we can find that the time is much shorter than that for the model considering only diffusion dynamics. However, this time is still longer than what is observed. Petschek reduced the reconnection region in order to resolve this discrepancy, resulting in the release time of $10^2\mbox{--}10^3 \ \mbox{s}$, which is closer to the observed one\cite{Petschek64}. As mentioned above, the theory of magnetic reconnection has been developed. However, the concept used in describing the time evolution of the magnetic field is that of diffusion. Hence, the question arises as to why diffusion produces a structural change of the magnetic field. Of course, we can often find a qualitative explanation by drawing analogy between magnetic field lines and rubber bands: the energy of short connected magnetic field lines is lower than that of extended magnetic field lines, as is shown in Fig.~\ref{fig:kakuta}. However, this analogy does not account for the aspect of diffusion dynamics. \begin{figure}[h] \begin{center} \includegraphics[scale=.32]{kakutaa.eps} \mbox{} \includegraphics[scale=.32]{kakutab.eps} \caption{\label{fig:kakuta}Illustration of magnetic reconnection by using an analogy between magnetic field lines and rubber bands. Enormous energy is stored in the extended magnetic field lines shown in (a). The occurrence of dissipation makes the magnetic force lines reconnect, as in (b), resulting in the release of stored energy.} \end{center} \end{figure} One of ways to determine that magnetic reconnection is caused by diffusion is to calculate the magnetic induction equation and other MHD equations numerically. However, such a calculation would be required to be done over a complex computer program using an expensive, high-specification machine. In this paper, we demonstrate how the magnetic field reconnects not through such an advanced technical process, but through a method that can be easily understood by university students who are familiar with diffusion phenomena. This paper is organized as follows. First, we derive the equations governing the time evolution of magnetic and velocity fields and define their initial and boundary conditions in Sec.~\ref{FE}. Next, the diffusion phenomena are revisited in Sec.~\ref{diff}, after which we demonstrate how the magnetic field lines are reconnected using the basic knowledge of diffusion in Sec.~\ref{main}. Finally, in Sec.~\ref{CR}, we conclude this paper and indicate the equality of our comprehension and the common qualitative explanation.
|
\label{CR} We have clarified how the diffusion gives rise to the magnetic reconnection. The magnetic induction equation has two components. First, the $x$ component, $B_x$, diffuses, and then, the $y$ component, $B_y$, which is initially zero, is generated by the time evolution of $B_x$. For the magnetic field lines to be reconnected correctly, the initial $B_x$ must be concave (convex) at $y>0$ ($y<0$) and must be saturated for $y$ far from the $x$ axis to such an extent as $1/a$ as in this paper; this behavior would indicate that the current density is localized around the $x$ axis, i.e., the current sheet. This current sheet is also necessary in the qualitative explanation. If an anomalous resistivity occurs, the current dissipates and the magnetic field starts to melt. As a result, enormous energy breaks out through the Joule dissipation and magnetic reconnection is generated. The common qualitative explanation of the magnetic reconnection using the current sheet is connected to our comprehension based on diffusion in this paper. Finally, we emphasize that if we adopt other concave (convex) functions of $y$ at $y>0$ ($y<0$), which are saturated as the initial magnetic field, magnetic reconnection can occur. \appendix
| 12
| 6
|
1206.5161
|
1206
|
1206.2924.txt
|
\pacs{11.27.+d, % Extended classical solutions; cosmic strings.. 98.70.Sa, % Cosmic rays (including sources, origin, acceleration, % and interactions) 98.80.Cq % Particle- and field-theory models of the early % universe (including cosmic strings...) } We study massive particle radiation from cosmic string kinks, and its observability in extremely high energy neutrinos. In particular, we consider the emission of moduli --- weakly coupled scalar particles predicted in supersymmetric theories --- from the kinks of cosmic string loops. Since kinks move at the speed of light on strings, moduli are emitted with large Lorentz factors, and eventually decay into many pions and neutrinos via hadronic cascades. The produced neutrino flux has energy $E \gtrsim 10^{11}~\rm{GeV}$, and is affected by oscillations and absorption (resonant and non-resonant). It is observable at upcoming neutrino telescopes such as JEM-EUSO, and the radio telescopes LOFAR and SKA, for a range of values of the string tension, and of the mass and coupling constant of the moduli.
|
Theories with spontaneous symmetry breaking usually have topologically non-trivial vacuum configurations. Depending on the topology of the vacuum after the symmetry breaking, stable relics called topological defects --- such as monopoles, strings or domain walls --- could be formed in the early universe \cite{Kibble:1976sj}. Strings can form if the vacuum manifold is not simply connected. %They were first studied in the context of superconductivity \cite{Abrikosov:1956sx} as flux tubes, and then, as solitonic solutions to abelian Higgs model \cite{Nielsen:1973cs}. Although monopoles and domain walls are generally problematic for cosmology, cosmic strings are compatible with the observed universe, provided that their tension is not too large (Sec.~\ref{general}; see, e.g., Refs.~\cite{VilenkinBook,Vachaspati:2006zz,Polchinski:2004ia,Copeland09,Copeland11} for reviews). Cosmic strings are predicted in grand unified theories (GUTs) and superstring theory, and their existence can be revealed through their effects on the cosmic microwave background (CMB), large scale structure and 21 cm line observations, and --- more directly --- by detecting their radiation, such as gravitational waves and cosmic rays. Since cosmic strings have GUT or superstring scale energy densities in their core, they can be significant sources of ultra high energy ($E \gtrsim 10^{11}$ GeV) cosmic rays \cite{Hill:1986mn,Bhattacharjee:1989vu,MacGibbon90,BOSV,Brandenberger:2009ia,Vachaspati10,BSV11}, either as isolated objects, or possibly in combination with other topological defects, like in monopole-string bound states \cite{Bhattacharjee:1994pk,Berezinsky:1997td,Berezinsky:1998ft,Berezinsky:1999az,BlancoPillado:2007zr}. Among the cosmic rays, neutrinos are especially interesting. Their weak coupling to matter makes them extremely penetrating, so they are the only form of radiation (together with gravitational waves) that can reach us from very early cosmological times, namely, all the way from redshift $z \sim 200$ (see Sec.~\ref{nuprop}). Moreover, in the spectral region of interest, $E \gta 10^{11}$ GeV, the \n\ sky is very quiet, since this region is beyond the range of \ns\ from even the most extreme hadron accelerators (gamma ray bursts, supernova remnants, active galactic nuclei, etc.). Therefore, even a low statistics \n\ signal beyond this energy would constitute a clean indication of a fundamentally different mechanism at play, such as a top-down scenario involving strings or other topological defects. Experimentally, the technologies to detect ultra high energy \ns\ are mature: they look for radio or acoustic signals produced by the \ns\ as they propagate in air, water/ice, or rock. After the successful experiences of ANITA \cite{anita}, FORTE \cite{Lehtinen:2003xv}, RICE \cite{rice} and NuMoon \cite{numoon} --- the latter using radio waves from the lunar regolith via the so called Askaryan effect \cite{Askarian} --- a new generation of experiments is being planned, that can probe \ns\ from cosmic strings with unprecedented sensitivity. Of these, the space based fluorescent light telescope JEM-EUSO \cite{jem}, and radio telescopes LOFAR \cite{lofar} and SKA \cite{ska} seem especially promising. One of the distinguishing effects of cosmic strings as cosmic ray emitters is that they can produce bursts from localized features called cusps and kinks (Sec.~\ref{general}), where ultrarelativistic velocities are reached. The radiation from cusps and kinks is very efficient, whereas the emission from cusp/kink-free string segments is exponentially suppressed. This enhanced emission has been studied in connection with gravitational waves \cite{Garfinkle:1987yw,Garfinkle:1988yi,DV}, and electromagnetic radiation \cite{Vilenkin:1986zz,Garfinkle:1987yw,Garfinkle:1988yi,BlancoPillado:2000xy} like gamma ray bursts \cite{Paczynski,Berezinsky:2001cp,Cheng:2010ae} and radio transients \cite{Vachaspati:2008su,Cai:2011bi,Cai12}, as well as \n\ bursts \cite{BOSV}. Among the several scenarios considered, there are a few that predict cosmic ray and \n\ fluxes at an observable level, e.g., Refs.~\cite{BOSV,Brandenberger:2009ia,Vachaspati10,BSV11}. One of these, Ref.~\cite{BSV11}, involves the decay of moduli --- massive scalar fields that arise in supersymmetric and superstring theories ---, that can have various masses and couplings to matter. Moduli with coupling stronger than gravity are fairly natural \cite{Goldberger00,Conlon07,Frey06,Burgess10,Cicoli11} and relatively unconstrained due to their very short lifetimes \cite{Sabancilar10}, compared to gravitationally coupled ones \cite{Damour97,Peloso03,Babichev05}. By decaying into hadrons, the moduli eventually generate a \n\ flux. In Ref.~\cite{BSV11} the emission of such moduli from string cusps, and the corresponding \n\ flux were discussed. In this paper, we elaborate on the theme of moduli-mediated \n\ production from strings, and study modulus emission from kinks. We show that the emission from kinks is very efficient, and is the dominant energy loss mechanism for the cosmic string loops for a wide range of the parameters. We calculate the \n\ flux expected at Earth after a number of propagation effects, mainly absorption due to resonant ($Z^0$ resonance channel) and non resonant \n-\n\ scattering. We find that the flux might be observable at near future surveys, JEM-EUSO, LOFAR and SKA, depending on the parameters. The structure of the paper is as follows. After discussing some generalities on strings and kinks in Sec.~\ref{general}, the modulus emission from a cosmic string kink is calculated in Sec.~\ref{kinkrad}. In Sec.~\ref{particleprod}, we discuss the decay of moduli, the properties of the hadronic cascade initiated by their decay into gluons, and propagation of extremely high energy neutrinos in the universe. In Sec.~\ref{nuflux}, estimates are given for the kink event rate, the neutrino flux, and its detectability by the existing and future neutrino detectors. We also discuss the constraint from high energy gamma ray observations. Finally, in Sec.~\ref{disc}, we give our conclusions. %=====================================================================================
|
\label{disc} Cosmic strings loops form as a result of reconnection of long strings, and self-intersection of large loops. Kinks arise naturally as a result of these processes. We studied how kinks can radiate moduli, particles that arise in the supersymmetric models of particle physics, and that can have various masses and couplings to matter. The decay of moduli into pions via hadronic cascades produces a flux of neutrinos, which can be observable depending on the parameters. Specifically, we considered the string tension $G \mu$, the modulus coupling constant $\alpha$ and mass $m$ as free parameters, and showed that neutrinos with energies $E \gtrsim 10^{11}$ GeV can be easily produced by cosmic string loops via this mechanism, with flux \beq E^{2} J_{\nu}(E) \sim 1.7 \times 10^{-7}\, \frac{\mu_{-17}^{3/7}\, \bar{\alpha}^{2}\, E_{11}^{1/7}}{p\, (\Gamma/50)^{11/7}}\,~\rm{\frac{GeV}{cm^{2}\,s\,sr}}. \eeq The hadronic cascade stops producing pions at the modulus rest frame energy of order $\epsilon \sim 1$ GeV. In the rest frame of the loop, this energy is boosted by the Lorentz factor $\gamma$, so that the the minimum observed energy of the neutrinos is: \beq E_{\rm{min}} (z) \sim 4.6 \times 10^{14} \frac{(\Gamma/50)^{1/2} \mu_{-17}^{1/2} m_{4}^{1/2} \epsilon_{\rm{GeV}}} {(1+z)^{7/4}}~\rm{GeV}. \eeq The \n\ flux is shown in Fig.~\ref{detect} for representative sets of parameters; the termination of the flux at $E_{\min}$ appears clearly. The figure also gives the flux sensitivity of various experiments, showing that the predicted flux is within reach for the next generation neutrino detectors such as JEM-EUSO, LOFAR and SKA. A distinctive feature of radiation from cosmic string kinks is that particles are emitted in a fan-like pattern, confined into a narrow ribbon, hence bursts from individual kinks can possibly be identified by timing and directional coincidence. In Eq.~(\ref{Nburst}), we estimated the number of neutrinos emitted by a kink, and the corresponding number of events in a detector of a given effective area. We found that, for the fiducial values of the parameters used in our analysis, multiple neutrinos can be seen in the field of view of the detector. If ultra high energy \ns\ are observed at future experiments, what would be possible to learn? Top-down mechanisms would offer natural explanations, and, among those, cosmic strings would be a favored candidate. Even in the framework of cosmic strings, however, data analyses will necessarily be model-dependent, and various models would have to be considered. Our scenario involving moduli is a possibility among many, and other intermediate states leading to \n\ production are possible, e.g., modulus emission from string cusps \cite{BSV11} and heavy scalar particle emission from cusps of superconducting strings \cite{BOSV}. Another possible generation mechanism of extremely high energy neutrinos could be the KK mode emission from cusps and kinks of cosmic F- and D-strings. The emission of KK modes of gravitons from cusps was studied in Refs.~\cite{DufauxKK1,DufauxKK2}, and various cosmological constraints have been put on the cosmic superstring tension. Depending on the parameters, observable neutrino fluxes might be produced by this mechanism as well. A discrimination between different models will require the combination of complementary data, probably the detection of gravitational wave/electromagnetic counterparts of neutrino signals \cite{jointGWnu,jointGWem}. The identification of point-like sources of extremely energetic \ns\ (bursts) would favor cosmic string kinks or cusps as sources, a hypothesis that would be substantiated further by the observation of accompanying gravitational wave and/or gamma ray bursts. To distinguish between kinks and cusps could be possible since the event rate is larger for kinks for the given values of the parameters. In addition to a possible discovery of topological defects, detecting a flux of ultra high energy \ns\ might reveal new pieces of the still incomplete puzzle of \n\ physics. Most interestingly, if the data show a $Z^0$ resonance dip, we might gather information on the \n\ mass and have another, perhaps more direct, evidence of the existence of the cosmological relic \ns. The information on the \n\ mass might be especially important if at least one \n\ is light enough to evade a direct mass measurement in the laboratory. It is important to consider, however, that the extraction of any information from data would be complicated by many theoretical uncertainties. Let us comment on the uncertainties and simplifying assumptions of our calculation. First of all, we worked in a flat matter dominated universe, and ignored the recent accelerated expansion period of the universe, whose effect can be at most about a factor of a few in our final estimates. We also approximated the neutrino fragmentation function for the moduli decays as $dN/dE \propto E^{-n}$, and used $n=2$, whereas the numerical calculations yield $n \approx 1.9$ \cite{DGLAP}. In our estimates we take into account the reconnection probability $p$. For cosmic strings of superstring theory, namely, F- and D-strings \cite{Jackson04}, $p<<1$, whereas for ordinary field theory strings $p=1$. The flux, event rate and the chance of getting neutrino bursts is expected to be enhanced for cosmic superstrings with $p\lesssim 1$, compared to ordinary cosmic strings. We ignored the backreaction of modulus emission from kinks on the evolution of kinks. Since the total power from a kink is only logarithmically divergent [see Eq.~(\ref{Ptot})], the effect of radiation is expected to smooth out the sharpness of a kink slowly. Finally, our treatment of the \n\ absorption due to resonant scattering on the \n\ backround is limited to relatively large masses, $m_{\nu} \gta 0.1$ eV, for which thermal effects on the background are negligible. The generalization to include these effects is forthcoming \cite{usprep}. %=======================================================================================
| 12
| 6
|
1206.2924
|
1206
|
1206.0744_arXiv.txt
|
Double white dwarf binaries in the Galaxy dominate the gravitational wave sky and would be detectable for an instrument such as LISA. Most studies have calculated the expected gravitational wave signal under the assumption that the binary white dwarf system can be represented by two point masses in orbit. We discuss the accuracy of this approximation for real astrophysical systems. For non-relativistic binaries in circular orbit the gravitational wave signal can easily be calculated. We show that for these systems the point mass approximation is completely justified when the individual stars are axisymmetric irrespective of their size. We find that the signal obtained from Smoothed-Particle Hydrodynamics simulations of tidally deformed, Roche-lobe filling white dwarfs, including one case when an accretion disc is present, is consistent with the point mass approximation. The difference is typically at the level of one per cent or less in realistic cases, yielding small errors in the inferred parameters of the binaries.
|
At low frequencies (mHz), millions of double white dwarf binaries in the Galaxy are expected to dominate the gravitational wave (GW) sky. At the lowest frequencies they form an unresolved foreground, while at frequencies above several mHz, thousands of sources would be individually detectable for an instrument such as the Laser Interferometer Space Antenna, LISA (\citealt{e1}, \citealt{l1}, \citealt{h1}, \citealt{nyp01, nyp03}) or eLISA/NGO (\citealt{eLISA}). These binaries come in two flavours: detached systems and semi-detached (mass-transferring) systems that are know as AM CVn systems (see \citealt{solheim2010}, \citealt{marsh2011} for reviews). Several known binaries should be detected by LISA within the first weeks of operation and are known as verification binaries (\citealt{sv2006}, \citealt{rgb+07, roelofs2010}, \citealt{brown2011}). By measuring their gravitational wave amplitude and frequency (evolution), the type of astrophysical source and its parameters can be determined (e.g. \citealt{cutler1994}, \citealt{littenberg2011}, \citealt{blaut2011}). In all these calculations the gravitational wave signal was determined under the assumption that the binary white dwarf system can be represented by two point masses in orbit, even for the tidally deformed stars in semi-detached binaries. The goal of this study is to determine the accuracy of this assumption. In section \ref{section gravwaves} we will discuss our method of calculating the GW signal. In section \ref{section baloons} we will give an algebraic view on the assumption of using point masses and in section \ref{section SPH} we will calculate the GW signal from smoothed-particle hydrodynamics (SPH) simulations of AM CVn stars. In section \ref{section conclusions} we discuss the conclusions of this study.
|
We have calculated the deviation of the gravitational wave signal of finite size and non-spherical white dwarfs to that of point masses, which is usually assumed. For any finite size, non-relativistic axisymmetric body in circular orbit, the result is exactly the same. For semi-detached white dwarf binaries, we find that an accretion disk of reasonable mass changes the gravitational wave signal at the level of $10^{-4}$ to $10^{-3}$, small but still significantly larger than errors due to the neglect of post-Newtonian corrections in the calculation of the signal. Deformations due to filling the Roche lobe of semi-detached binaries increase for mass ratios closer to unity for fixed donor mass and are distinctly stronger at fixed mass ratio for lower mass donors and can in the most extreme cases be of order 1 per cent. This in principle will change the frequency evolution and the accuracy with which the parameters can be determined (e.g. \citealt{blanchet1995}, \citealt{cutler1994}), so calculations in which accuracies of better than one per cent are needed should take the finite size into account. Also, the different strength of the gravitational wave angular momentum losses will affect the mass transfer rate and stability of the mass transfer in semi-detached systems (e.g. \citealt{marsh04}). However, the level of deviation we found shows that the calculations presented in the literature on the expected signals of (verification) binaries for LISA, are essentially unaffected: for monochromatic sources, the amplitude of the signal will be slightly different, but indistinguishable from sources with slightly higher (chirp) masses and/or smaller distances, properties that are much more uncertain than one per cent. For systems with measurable period derivatives, the degeneracy between (chirp) mass and distance is broken, but the deviations we found can still only be detected if there is independent extremely accurate measurement of the (chirp) mass of the binary. For a LISA-like detector such independent mass estimates, if available, are typically accurate at the 10 per cent level at best (e.g. \citealt{littenberg2011}). The detailed evolution (and possible merger of the system) will be affected as well, but the calculations of this phase (e.g. \citealt{Dan2012}) already take the finite size of the stars into account. We therefore conclude that in the majority of cases, the use of the point mass approximation is well justified.
| 12
| 6
|
1206.0744
|
1206
|
1206.0602_arXiv.txt
|
The \textit{LOw Frequency ARray} (LOFAR) is a next-generation radio telescope which uses thousands of stationary dipoles to observe celestial phenomena. These dipoles are grouped in various `stations' which are centred on the Netherlands with additional `stations' across Europe. The telescope is designed to operate at frequencies from 10 to 240\,MHz with very large fractional bandwidths (25-100\%). Several `beam-formed' observing modes are now operational and the system is designed to output data with high time and frequency resolution, which are highly configurable. This makes LOFAR eminently suited for dynamic spectrum measurements with applications in solar and planetary physics. In this paper we describe progress in developing automated data analysis routines to compute dynamic spectra from LOFAR time-frequency data, including correction for the antenna response across the radio frequency pass-band and mitigation of terrestrial radio-frequency interference (RFI). We apply these data routines to observations of interplanetary scintillation (IPS), commonly used to infer solar wind velocity and density information, and present initial science results.
|
\label{sec:intro} The observation of interplanetary scintillation (IPS) -- the scintillation of compact radio sources due to density variations in the solar wind \citep{Hewishetal:1964} -- is an important tool for observing the solar wind. Observations of IPS allow the solar wind speed to be inferred over all heliographic latitudes and a wide range of elongations from the Sun \citep[e.g.][]{DennisonHewish:1967}, giving a global perspective to point measurements from spacecraft. The sensitivity of IPS to small, turbulent-scale, density variations also complements the larger-scale sensitivities of white-light observations from coronagraphs. Advances in the study of IPS over the last decade or more allow such observations to be used to calculate three-dimensional (3D) sky maps of solar wind electron density and speed \citep[e.g.][]{Asaietal:1998, Jacksonetal:1998, Kojimaetal:1998}. These maps and other detailed analyses of observations of IPS \citep[e.g.][]{Fallowsetal:2008a, Breenetal:2008, Bisietal:2007} are increasingly recognised as a valuable aid for tracking space weather events through the inner heliosphere \citep[e.g.][]{Jacksonetal:2010} and to the overall study of space weather prediction. The \textit{LOw-Frequency ARray} (LOFAR -- summarised fully in Section \ref{sec:lofar}) is a major new-generation radio telescope operating in the 10--240\,MHz frequency range. It consists of arrays of dipoles grouped into stations with a central `core' of stations in the Netherlands and, currently, eight international stations based in the UK, France, Germany and Sweden. Although designed principally to be used as a single array, it is also possible to use the stations individually making it suitable for studies of IPS. A particular advantage offered by LOFAR is the ability to observe with bandwidths of up to 48\,MHz with a high frequency resolution. This capability is both necessary to identify and eliminate radio frequency interference (RFI) and useful to create dynamic spectra of IPS data, a tool that could provide new insights into solar wind microstructure and has not been available to most prior IPS observing instruments. This paper describes the principle science of IPS, details how LOFAR may be used as an instrument to observe it, the principle advantages LOFAR can offer as such an instrument and the progress made in obtaining observations of IPS. We discuss in particular on-going efforts to develop an automated Python-based software `pipeline' to produce relevant products from raw LOFAR data. Over time these tools will be developed to enable the study of other similar phenomena, such as flare stars, planetary atmospheres and solar radio bursts. The paper is laid out as follows: The study of IPS is summarised in Section \ref{sec:ips}; Section \ref{sec:lofar} describes the LOFAR radio telescope and details how it may be used for IPS; the current state of a dynamic spectrum data pipeline developed for LOFAR observations is detailed in Section \ref{sec:rfi} and then some initial IPS results are presented in Section \ref{sec:results}.
|
\label{sec:conclusions} The dynamic spectrum results obtained so far hint at a wealth of new information on solar wind micro-structure and turbulence. Previous studies are few, but \citet{ColeSlee:1980} did observe a dynamic spectrum over the frequency range 280--520\,MHz of IPS seen in an observation of 3C273. This study showed a curve in the scintillation maxima over the frequency band, later attributed to refraction due to large-scale components of a Kolmogorov turbulence regime in the solar wind \citep{ColesFilice:1984}. Such a curve is not seen in LOFAR observations of IPS taken so far, most likely because the 40--48\,MHz bandwidth used is not large enough for it to be seen clearly. The level of scintillation is known to vary with both distance from the Sun and observing frequency \citep[e.g.][]{Coles:1978}; within the `weak' scattering regime, lower observing frequencies will exhibit stronger scintillation at the same solar elongation. The observation of 3C48 shown in Figure \ref{fig:dynspec} illustrates this nicely, with the scintillation being stronger at the lower frequencies. The power spectra shown in Figure \ref{fig:spectra} are consistent with typical IPS spectra, but also show some inconsistencies. The excess power evident at higher spectral frequencies in the observation of 3C48 and the inconsistency in the observation of 3C84 when compared to a simultaneous observation from a known antenna are particular points of concern. It is possible that at least some of these inconsistencies are due to the high LOFAR observing frequencies used: It is known that a grating lobe of the main station beam (an artifact of the use of an array of dipoles) can be present above the horizon for frequencies in the high-band of LOFAR (Ger de Bruyn, private communication, 2011), potentially causing issues with excess noise. This is not well-characterised yet making predictions of whether it will or will not be above the horizon for a particular observation difficult. It is clear from all the observations of IPS taken so far that the excess power noted in the observation of 3C48 in Figure \ref{fig:spectra} is apparent in many observations using the HBA, but not in all. Restricting the sub-bands used in creating the time series to those at the lower or higher ends of the observing band may occasionally make a difference but, again, not in every case. The correlation functions shown in Figure \ref{fig:correlations} indicate the presence of slow solar wind streams in the lines of sight. This is consistent with the low heliographic latitudes of both these observations. The cross-correlation function of the 3C279 observation shows a `negative lobe' at zero time-lag. This is often associated with the presence of a Coronal Mass Ejection in the line of sight. A slow CME was observed to launch from the Sun late on 14 November 2011 and was predicted to pass close to the Earth on 18-19 November 2011. An initial check of coronagraph data indicate that this CME was launched in a direction that could cross the line of sight of the IPS observation and the observation timing makes it a promising candidate. However a full geometrical analysis will be required to confirm whether or not this CME is seen in these IPS data. In conclusion, these observations show a high degree of promise, but also reveal that some issues remain. This is to be expected from an instrument which is still undergoing commissioning. \begin{acks} LOFAR, the \textit{Low Frequency Array} designed and constructed by ASTRON, has facilities in several countries, that are owned by various parties (each with their own funding sources), and that are collectively operated by the International LOFAR Telescope (ILT) foundation under a joint scientific policy. The authors thank the director and staff of EISCAT for the ESR data used in this study. EISCAT is funded by the research councils of Norway, Sweden, Finland, Japan, China, the United Kingdom and Germany. Two of us (RAF and MMB) were funded by the UK Science and Technology Facilities Council during the course of this work. \end{acks}
| 12
| 6
|
1206.0602
|
1206
|
1206.3571_arXiv.txt
|
Satellite galaxies in groups and clusters are more likely to have low star formation rates (SFR) and lie on the `red-sequence' than central (`field') galaxies. Using galaxy group/cluster catalogs from the Sloan Digital Sky Survey Data Release 7, together with a high-resolution, cosmological \tit{N}-body simulation to track satellite orbits, we examine the star formation histories and quenching timescales of satellites of $\mstar > 5 \times 10 ^ {9} \msun$ at $z \approx 0$. We first explore satellite infall histories: group preprocessing and ejected orbits are critical aspects of satellite evolution, and properly accounting for these, satellite infall typically occurred at $z \sim 0.5$, or $\sim 5 \gyr$ ago. To obtain accurate initial conditions for the SFRs of satellites at their time of first infall, we construct an empirical parametrization for the evolution of central galaxy SFRs and quiescent fractions. With this, we constrain the importance and efficiency of satellite quenching as a function of satellite and host halo mass, finding that satellite quenching is the dominant process for building up all quiescent galaxies at $\mstar < 10 ^ {10} \msun$. We then constrain satellite star formation histories, finding a `delayed-then-rapid' quenching scenario: satellite SFRs evolve unaffected for $2 - 4 \gyr$ after infall, after which star formation quenches rapidly, with an e-folding time of $< 0.8 \gyr$. These quenching timescales are shorter for more massive satellites but do not depend on host halo mass: the observed increase in satellite quiescent fraction with halo mass arises simply because of satellites quenching in a lower mass group prior to infall (group preprocessing), which is responsible for up to half of quenched satellites in massive clusters. Because of the long time delay before quenching starts, satellites experience significant stellar mass growth after infall, nearly identical to central galaxies. This fact provides key physical insight into the subhalo abundance matching method.
|
Observations have long shown that galaxies in denser regions are more likely to have low star formation rates (SFR), lie on the red sequence, and exhibit more evolved (elliptical) morphologies than similar mass galaxies in less dense regions, from massive galaxies in clusters \citep{Oem74, DavGel76, Dre80, DreGun83, PosGel84, BalMorYee97, PogSmaDre99} to the lowest mass satellites in the Local Group \citep{Mat98}. Large-scale galaxy surveys, such as the Sloan Digital Sky Survey \citep [SDSS;][]{SDSS}, have enabled detailed examinations of the correlations between these galaxy properties and their environment at $z \approx 0$ \citep[see][for a recent review]{BlaMou09}. Several such early works showed that galaxy SFR/color depends on small-scale ($\lesssim 1 \mpc$) environment, with little-to-no additional dependence on larger scale environment \citep{HogBlaBri04, KauWhiHec04, BlaEisHog05}. More physically, this environmental dependence has been shown to result from satellite galaxies and the properties of their host dark matter halo \citep{WeivdBYan06a, BlaBer07, WilZibBud10, TinWetCon11}, where `satellite' galaxies are all those that are not the massive `central' galaxy at the core of the host halo. These results are physically meaningful, given that the virial radius corresponds to a physical transition from the low-density `field' environment to a high-density, virialized region. After a satellite falls into a host halo, the strong gravitational tidal forces prevent the satellite's (sub)halo from accreting dark matter and also strip mass from the subhalo from the outside-in \citep[e.g.,][]{DekDevHet03, DieKuhMad07}. Additionally, if the host halo is massive enough to host a stable, virial accretion shock \citep{DekBir06}, then its thermalized gas also can heat and strip any extended gas in the subhalo \citep{BalNavMor00, KawMul07, McCFreFon08}. Therefore, satellites (eventually) experience reduced gas cooling/accretion rates onto their disc after infall, a phenomenon known as `strangulation' or `starvation' \citep{LarTinCal80}. More drastically, in the extreme case of both high gas density and satellite velocity, ram-pressure can strip cold gas directly from the disc \citep{GunGot72, AbaMooBow99, ChuvGoKen09}. The dense collection of galaxies in a host halo also allows for the possibility of strong gravitational interactions with neighboring galaxies, known as `harassment' \citep{FarSha81, MooLakKat98}, and satellites can merge with one another \citep{MakHut97, AngLacBau09, WetCohWhi09a, WetCohWhi09b, WhiCohSmi10, Coh12}. All of these mechanisms are expected to play some role in quenching satellite star formation, though their importance, particularly as a function of host halo mass, remains in debate. To constrain satellite quenching processes, many works have examined the SFRs/colors of satellites in groups/clusters in detail at $z \approx 0$ \citep[e.g.,][]{BalNavMor00, EllLinYee01, DePColPea04, WeivdBYan06a, BlaBer07, KimSomYi09, vdBAquYan08, HanSheWec09, KimSomYi09, PasvdBMo09, vdLWilKau10, PreBalJam11, PenLilRen11, WetTinCon12a, WooDekFab13}. In general, these works found that the fraction of satellite that are quiescent/red, $\fsatq$, depends primarily, and independently, on three quantities: $\fsatq$ (1) increases with satellite mass, (2) increases with the mass of the host halo, and (3) increases toward halo center. Trend (1) is caused, at least partially, by the underlying dependence on stellar mass set by central galaxies prior to infall. Trend (2) is sometimes interpreted as satellites being quenched more rapidly in more massive host halos, but the hierarchical nature of halo growth, namely, the possibility of quenching as a satellite in a lower mass halo prior falling into a more massive halo (`group preprocessing') complicates this interpretation \citep[e.g.,][]{ZabMul98, McGBalBow09}. Finally, trend (3) implies an evolutionary trend, because a satellite's halo-centric radius negatively correlates with its time since infall \citep[e.g.,][]{GaoWhiJen04, DeLWeiPog12}. Similar satellite trends persists out to at least $z \sim 1$ \citep[e.g.,][]{CucIvoMar06, CooNewCoi07, GerNewFab07, TraSaiMou09, PenLilKov10, McGBalWil11, GeoLeaBun11, MuzWilYee12}. Several works have gone beyond simple satellite SFR/color cuts to examine observationally the nature of the full SFR/color distribution. These works have shown that the color \citep{BalBalNic04, Ski09} and SFR \citep{BalEkeMil04, McGBalWil11, PenLilRen11, WetTinCon12a, WooDekFab13, WijHopBro12} distribution of galaxies is strongly bimodal across all environmental/halo regimes, and the SFR/color of actively star-forming/blue galaxies does not vary with any environmental measure. As noted in many of the above works, these results imply that the environmental process(es) takes considerable time (several Gyrs) to affect satellite SFR. In this paper, we seek to use the aforementioned observational trends, which we presented in detail in \citet{WetTinCon12a}, to quantify---in a robust, statistical manner---the star formation histories and quenching timescales of satellite galaxies at $z = 0$ across a wide range of both satellite and host halo masses. Understanding satellite quenching mechanisms and the timescales over which they operate is important for elucidating the physical processes that occur in groups and clusters, but also for a comprehensive understanding of galaxy evolution overall, given that satellites constitute $\sim 1 / 3$ of all low-mass galaxies \citep[e.g.,][]{YanMovdB07}. Satellite galaxies also provide unique laboratories for examining gas depletion and its relation to star formation because, unlike central galaxies, satellites are thought not to accrete gas from the field after infall. Furthermore, because satellites are significantly more likely to lie on the red sequence, many methods for identifying galaxy groups/clusters rely on selecting red-sequence galaxies \citep[e.g.,][]{GlaYee00, KoeMcKAnn07a}, so a detailed understanding of the systematics of these methods requires characterizing the timescale over which satellites migrate onto the red sequence after infall and how this timescale depend on host halo mass and redshift. Many works have investigated satellite SFR evolution and quenching through the use of semi-analytic models (SAMs) applied to cosmological \tit{N}-body simulations. In one early example, \citet{BalNavMor00} modeled satellite SFR as declining exponentially after infall on a cold gas consumption timescale of a few Gyrs to account for radial gradients of average SFR and color in clusters. Many SAMs assumed that a satellite subhalo's hot gas is stripped instantaneously as it passes within the host halo's virial shock, but this scenario quenches star formation too rapidly; only models that remove/strip gas more gradually produce realistic quiescent fractions \citep{WeivdBYan06b, FonBowMcC08, KanvdB08, BooBen10}, but, in general they have difficulty in correctly reproducing the full SFR distribution. Though, \citet{WeiKauvdL10} recently implemented a modification of the SAM of \citet{DeLBla07} in which the diffuse gas around a satellite galaxy is stripped at the same rate as its host dark matter subhalo (10 - 20\% loss per Gyr), showing that this modification produces a satellite SSFR distribution that broadly is in agreement with observations. However, understanding the results of SAMs is complicated by the fact that they do not completely accurately model the evolution of central galaxies, so they do not provide fully accurate initial conditions at infall for satellites. Relatedly, a few works have examined satellite SFR evolution in cosmological hydrodynamic simulations of galaxy groups \citep[e.g.,][]{FelCarMay11}, arguing that quenching occurs largely through the lack of gas accretion and (to a lesser degree) gas stripping after infall. Instead of attempting fully to forward-model all of the relevant physical processes for satellites, our approach is to parametrize satellite star formation histories and constrain their quenching timescales in as much of an empirical manner as possible. We start with detailed measurements of the SFRs of satellites at $z = 0$ from our SDSS group catalog that we presented in \citet{WetTinCon12a} and review in \S\ref{sec:method}. We also describe our cosmological $N$-body simulation, which we use to create a mock group catalog to compare our models to observations robustly. With this simulation, we explore the infall times of satellites in \S\ref{sec:infall-time}. We then develop an accurate, empirical parametrization for the initial SFRs of satellites at their time of infall in \S\ref{sec:sfr_at_infall}. Having accurate initial SFRs of satellites and measurements of their final SFRs at $z = 0$, we examine the importance and efficiency of satellite quenching in \S\ref{sec:quench_importance_efficiency} and their star formation histories and quenching timescales in \S\ref{sec:sfr-evol_sat}. With this, we then examine where satellites were when they quenched in \S\ref{sec:where_quench} and their stellar mass growth in \S\ref{sec:m-star_growth}. Finally, we discuss the implications of our results for subhalo abundance matching in Appendix \ref{sec:mass_growth_sham}. This paper represents the third in a series of four. In \citet{TinWetCon11}, hereafter \citetalias{TinWetCon11}, we described our SDSS galaxy sample, presented our method for identifying galaxy groups/clusters, and showed that central and satellite galaxy quiescent fractions are essentially independent of the large-scale environment beyond their host halo. In \citet{WetTinCon12a}, hereafter \citetalias{WetTinCon12a}, we used our SDSS group catalog to examine in detail the SFR distribution of satellite galaxies and its dependence on stellar mass, host halo mass, and halo-centric radius, finding that the SFR distribution is strongly bimodal in all regimes. Based on this, we argued that satellite star formation must evolve in the same manner as central galaxies for several Gyrs after infall, but that once satellite quenching starts, it occurs rapidly. In \citet{WetTinCon13a}, hereafter \citetalias{WetTinCon13a}, we will examine quenching in galaxies \tit{near} groups and clusters, focusing on `ejected' satellites that passed within a more massive host halo but have orbited beyond the virial radius. We will show that these ejected satellites can explain essentially all trends for star formation quenching in galaxies beyond the the virial radius of groups/clusters. Finally, in \citet{WetTinCon13b}, hereafter \citetalias{WetTinCon13b}, we will use the detailed orbital histories from our simulation to constrain the physical mechanisms responsible for satellite quenching. For clarity, we outline some nomenclature. We refer to galaxies as `quiescent' in an observational sense: having low SFR but without regard to how or when SFR faded. By contrast, we refer to satellite `quenching' in our models as the physical process of SFR fading rapidly below the quiescence threshold, under the ansatz that once a satellite is quenched it remains so indefinitely. Our galaxy group catalog refers to `group' in a general sense, as a set of galaxies that occupy a single host halo, regardless of its mass, and we will use `(host) halo' as a more general term for group or cluster. Finally, we cite all masses using $h = 0.7$ for the Hubble parameter.
|
\label{sec:summary} \subsection{Summary} \begin{figure} \centering \includegraphics[width = 0.99 \columnwidth]{fig12.pdf} \caption{ Summary diagram of satellite galaxy SFR evolution, as given by equation (\ref{eq:sfr-evol_sat}), highlighting the `delayed-then-rapid' quenching scenario. The physical process(es) responsible for quenching star formation in a satellite begins after it first falls into another host halo, regardless of the host halo's mass. However, it takes considerable time for SFR to be affected: satellite SFR evolves after infall in the same manner as central galaxies for a delay time $\tqdelay = 2 - 4 \gyr$ (red curve), depending on stellar mass. Then, the satellite's star formation starts to be quenched, and SFR fades rapidly, with an e-folding time $\tauqfade = 0.2 - 0.8 \gyr$, also depending on stellar mass. Less massive satellites have longer $\tqdelay$ and $\tauqfade$, but neither timescale dependends on the mass of the host halo. Because of the long $\tqdelay$, satellite stellar mass growth via star formation is nearly equal to that of central galaxies. Because of hierarchical halo growth, many satellites in massive host halos were quenched as a satellite in a lower mass halo prior to infall. For comparison, blue curve shows gradual SFR fading for central galaxies, with characteristic fading time $\taucen = 2 - 4 \gyr$ from equation (\ref{eq:sfr-evol_cen}). } \label{fig:sfr-evol_diagram} \end{figure} Using a galaxy group/cluster catalog from SDSS Data Release 7, together with a cosmological \tit{N}-body simulation to track satellite orbits, we examined in detail the star formation histories of satellite galaxies at $z \approx 0$, focusing on their times since infall, quenching timescales, and stellar mass growth after infall. Applying the same group-finding algorithm to our simulation as we used in SDSS allows us to make robust comparisons of model results to observations. To obtain accurate initial conditions for the SFRs of satellites at their time of first infall, we constructed an empirically based, statistical parametrization for the evolution of central galaxy SFRs out to $z = 1$; this is critical for the accuracy of our results because, at fixed stellar mass, the quiescent fraction for central galaxies more than doubles from $z = 1$ to 0. Our primary result is that, at least on average, satellite SFR evolves via a `delayed-then-rapid' quenching scenario: satellite SFR remained unaffected for several Gyrs after first infall, after which quenching occurs rapidly, as Fig.~\ref{fig:sfr-evol_diagram} summarizes. In more detail, our main results are: \tit{Infall as part of a group and ejection beyond $\rvir$ are important aspects of satellite evolution.} Fewer than half of satellites in massive clusters fell in directly from the field; the rest fell in as a satellite in another host halo or experienced secondary infall after becoming ejected. Satellites at $z = 0$ experienced their first infall typically at $z \sim 0.5$, or $\sim 5 \gyr$ ago, with a broad tail out to $z \ge 1$. Less massive satellites and those in more massive host halos fell in earlier. \tit{Satellite quenching is a critical process of galaxy evolution.} Satellite quenching always dominates the production of quiescent satellites. Moreover, satellite quenching is responsible for producing the majority of \tit{all} quiescent (red-sequence) galaxies at $\mstar < 10 ^ {10} \msun$ by $z = 0$. \tit{Satellite quenching is delayed-then-rapid.} Based on observations, we argued that the process(es) responsible for quenching satellites begins after \tit{first} infall into any other host halo. As constrained by the satellite SSFR distribution at $z = 0$, satellites, at least on average, then remain actively star-forming for $2 - 4 \gyr$ (depending on stellar mass) after first infall, unaffected by their host halo, before quenching starts. Once quenching has started, the e-folding time over which SFR fades is $< 0.8 \gyr$. \tit{Satellite quenching timescales are shorter at higher stellar mass but are independent of host halo mass.} More massive satellites start to be quenched more rapidly after infall, and once quenching has started, their SFRs fade more quickly. However, these quenching timescales do not depend on host halo mass. The observed increase in the satellite quiescent fraction with host halo mass arises because of the increased importance of group preprocessing and ejection/re-infall in more massive host halos. \tit{Group preprocessing plays a critical role in quenching satellites.} Half of low-mass ($\mstar < 10 ^ {10} \msun$) quiescent satellites in clusters started quenching in another host halo before falling into the cluster. Across all satellite masses, the fraction of quiescent satellites that were quenched within their current host halo is never more than $\sim 70\%$. \tit{Satellite quenching barely impacts stellar mass growth.} Because satellite quenching is so delayed, low-mass satellites have experienced considerable stellar mass growth via star formation since infall: satellites at $\mstar < 10 ^ {10} \msun$ are, on average, 50\% more massive than at infall. Moreover, the average amount of mass growth via star formation in satellite and central galaxies is identical to within 10\%. This provides key physical insight into the abundance matching technique for assigning stellar mass to subhalos, as outlined in Appendix \ref{sec:mass_growth_sham}. \subsection{Relation to satellite gas content} \label{sec:quench_time_gas} We first discuss the relation of our results to satellite gas content. Satellites provide unique laboratories for examining gas depletion and its relation to star formation because, unlike central galaxies, satellites' subhalos are thought not to accrete matter after infall: the strong gravitational tidal forces in the host halo both prevent a satellite's subhalo from accreting new matter and strip any existing subhalo matter, including gas, from the outside-in. Additionally, any thermalized gas in the host halo can heat and ram-pressure strip any extended subhalo gas, and in the extreme case of both high gas density and satellite velocity, ram-pressure can strip cold gas directly from the disc. We first discuss the implications of our $\tqdelay$ results, that is, that SFR in satellites evolves for $2 - 4 \gyr$ after infall, depending on stellar mass, \tit{unaffected by the host halo}. While this timescale may represent, to some degree, the statistical average of those of individual satellites, it is informative to consider in the context of gas depletion times. Given that star formation is fueled by cold, molecular gas \citep{WonBli02, BigLerWal08}, this means that a significant quantity of such gas must persist in a satellite's disc for that amount of time. One possibility is that a sufficient reservoir of cold gas was present in the disc at the time of infall. As a constraint, we compare our $\tqdelay$ times to observed cold gas depletion times, defined as $M_{\rm gas} / \sfr$. At $z = 0$, observed atomic gas depletion times in $\mstar > 10 ^ {10} \msun$ galaxies are $\sim 3 \gyr$, with large scatter but no systematic dependence on stellar mass or SFR \citep{SchCatKau10}. Incorporating the additional $\sim 30\%$ of the gas that is molecular \citep{SaiKauKra11}, the total gas depletion time would extend to $\sim 4 \gyr$. This timescale can be even longer to the extent that gas recycled from stellar mass loss fuels star formation \citep[e.g.,][]{LeiKra11}. If valid at higher redshift, this depletion time would be sufficient to accommodate our $\tqdelay$ values \tit{if} all of the atomic gas converts to stars. Observed gas ratios, $M_{\rm gas} / \mstar$, provide another constraint. In \S\ref{sec:m-star_growth}, we showed that satellites experience significant stellar mass growth via star formation. In particular, currently quiescent, low-mass satellites have more than doubled their stellar mass since infall, which requires a gas reservoir comparable in mass to their stars at the time of infall. Observations at $z = 0$ show that actively star-forming galaxies at $\mstar \sim 10 ^ {10} \msun$ have total cold gas masses that are $\sim 40\%$ of their stellar mass \citep{CatSchKau10, SaiKauKra11}, and that this gas ratio increases with decreasing stellar mass, being near unity for galaxies just below our mass threshold \citep[e.g.,][]{GehBlaMas06}. However, currently quiescent, low-mass satellites fell in at higher redshift (typically, $z = 0.5 - 1$), and if the total cold gas fraction increases with redshift at a rate suggested by observations of molecular gas in actively star-forming, massive ($\mstar > 10 ^ {10} \msun$) galaxies at $z = 0.4 - 1.4$ \citep{DadBouWal10, TacGenNer10, GeaSmaMor11}, then these satellites would have had enough cold gas in their disc to accommodate the significant stellar mass growth in Fig.~\ref{fig:m-star-growth_v_m-star}. Furthermore, the cold gas in the disc could be replenished for some time after infall if the most concentrated and tightly-bound component of the extended subhalo gas continues to cool/accrete onto the disc for several Gyrs before being stripped. X-ray observations show that roughly half of massive satellites in groups and clusters retain extended, hot gas halos, though truncated as compared with central (`field') galaxies \citep{SunDonVoi07, JelBinMul08}. Simulations also show retention of extended subhalo gas: \citet{McCFreFon08} found that satellite subhalos can retain a significant fraction ($\sim 30\%$) of their hot gas for several Gyrs after infall, while \citet{SimWeiDav09} and \citet{KerKatFar09} found that satellites continue to accrete significant gas onto their disc, though at a reduced rate compared with central galaxies. Adding this replenishment to what cold gas was already in the disc at infall, the total gas reservoir in/around satellites appears fully sufficient to fuel their extended SFR and stellar mass growth as demanded by $\tqdelay$. Finally, regarding $\tqdelay$, we note that the halo radius crossing time, given our virial definition, is $t_{\rm cross} = \rvir / \vvir = 2.7 (1 + z) ^ {-3 / 2} \gyr$, independent of host halo mass. More precisely, numerically integrating satellite orbits in an NFW potential (assuming energy and angular momentum conservation) using typical initial orbital parameters from \citet{Wet11} for satellites in host halos in our mass range, the average time from infall to first pericentric passage is somewhat shorter at $2 (1 + z) ^ {-3 / 2} \gyr$, independent of satellite mass. Thus, the onset of satellite quenching occurs near the time of pericentric passage for more massive satellites and $1 - 2 \gyr$ after pericentric passage for lower mass satellites, as we will explore in more detail in \citetalias{WetTinCon13b}. We next discuss the implications of our $\tauqfade$ results, that is, that once quenching has started, satellite SFR fades rapidly, with $\tauqfade$ being 0.8 to $0.2 \gyr$ from $\mstar = 5 \times 10 ^ 9$ to $2 \times 10 ^ {11} \msun$. A lack of star formation implies a lack of dense, molecular gas, so it is interesting to compare $\tauqfade$ to observed molecular gas depletion times, defined as $M_{\rm H_2} / \sfr$, for galaxies near the quenching threshold. \citet{SaiKauWan11} examined the molecular gas depletion times in a large sample of SDSS galaxies, finding typically $\sim 1 \gyr$ at $\mstar = 10 ^ {10 - 11} \msun$. However, they found that the depletion time increases with decreasing SSFR, being $\sim 2 \gyr$ at $\ssfr \sim 10 ^ {-11} \yrinv$. Such a long depletion time in galaxies near the quenching threshold is difficult to understand in a scenario in which satellites simply use up their molecular gas, which may suggest that additional processes are at play in reducing the molecular gas density in satellites as they are quenched, possibly via tidal or ram-pressure stripping or internal feedback processes. However, note that the sample in \citeauthor{SaiKauWan11} is composed primarily of central (`field') galaxies, and it is unclear whether the significant gas reservoirs in these nearly quiescent galaxies result from them not fully depleting their cold gas while quenching, or by them accreting gas after they have quenched. If the latter holds, the cold gas properties of satellites as they are being quenched could be quite different. Several works have examined the gas content of satellites in the Virgo cluster, finding that while the atomic gas masses of Virgo satellites are significantly lower than for galaxies of the same mass in the field \citep{HuaHayGio12, SerOosMor12}, the molecular gas masses are quite similar \citep{KenYou86, YouBurDav11}. This result suggest that, while processes like tidal or ram-pressure stripping may play a role in removing atomic gas from the outer regions of satellites, they have little impact on the molecular gas that fuels star formation. Indeed, observations of ram-pressure stripping \citep{SunDonVoi07, ChuvGoKen09, AbrKenCrow11} typically show diffuse, atomic gas being stripped from the outer regions of the disc, while the dense, molecular gas towards the core survives intact, a phenomenon also seen in simulations \citep{TonBry09}. Overall, we conclude that the gas reservoir in/around satellites at their time of infall is sufficient to fuel their necessary star formation histories and stellar mass growth. Our result that satellite quenching can be parametrized simply by time since first infall, with no significant dependence on host halo mass, suggests that simple gas depletion (`strangulation') most naturally explains satellite quenching, though more work is needed to understand if gas self-depletion alone can account for our `delayed-then-rapid' quenching scenario with sufficiently short $\tauqfade$. Our $\tauqfade$ values are marginally-to-significantly shorter than observed molecular gas depletion times, particularly for galaxies near the quenching threshold, possibly suggesting that some process other than simple molecular gas depletion is at play, though it is not clear if external stripping processes can explain this. In \citetalias{WetTinCon13b}, we will examine these issues in more detail by developing physical models for satellite SFR evolution. \subsection{Comparison with other work} Our satellite quenching timescales are broadly consistent with previous works that parametrized the evolution of SFR in satellites and argued that it is affected over long ($2 - 3 \gyr$) timescales \citep[e.g.,][]{BalNavMor00, WanLiKau07, McGBalBow09, MahMamRay11}. Recently, \citet{DeLWeiPog12} examined the infall times of satellites in a SAM applied to the Millennium simulation, accounting for hierarchical halo growth and group preprocessing; comparing with observed satellite quiescent fractions as a function of host halo mass and halo-centric radius, they argued that satellites take $\sim 5 - 7 \gyr$ to quench after falling into halos $> 10 ^ {13} \msun$. However, these previous works generally only used observed quiescent/red fractions to constrain a single, overall quenching timescale, as in our Fig.~\ref{fig:quench-time_v_m-star}. A significant aspect of our approach is using the overall SSFR distribution to constrain satellite SFR evolution more completely, through which we have shown that satellites experience a `delayed-then-rapid' quenching scenario, which is not possible measuring just quiescent fractions. Furthermore, our use of empirically based satellite initial SFRs and a mock simulation group catalog to compare robustly with observations allows us to constrain these timescales empirically and robustly. Our results place strong constraints on semi-analytic approaches to modeling the physics of satellite SFR evolution \citep[e.g.,][]{FonBowMcC08, KanvdB08, WeiKauvdL10, KimYiKho11}. Our results suggest that a successful physical model would allow satellite SFR to evolve, environmentally unaffected, for $2 - 4 \gyr$ (depending on stellar mass) after infall, possibly through continued accretion/cooling of extended subhalo gas. Our results also suggest that one does not need to impose any explicit dependence on host halo mass to this process. We emphasize that our quenching timescales are valid for satellites at $z \sim 0$, given that our approach is sensitive to satellite SFR evolution within the last $\sim 4 \gyr$, and it is not clear that these timescales remain fixed at higher redshift. We have checked this in our framework by applying our quenching timescales from $z = 0$ to satellites at higher redshift and comparing with the observed evolution in the quiescent fraction for all galaxies from Fig.~\ref{fig:qu.frac_v_z}. While our model does agree within observational uncertainties at $z \lesssim 0.3$, we find that this approach quenches too few galaxies at higher $z$. One one hand, this discrepancy could be interpreted as evidence against the accuracy of our 'delayed-then-rapid' quenching scenario. Alternately, satellite quenching times simply may be shorter at higher redshift. Using halo occupation modeling of galaxy spatial clustering measurements, \citet{TinWet10} showed that the satellite quiescent fraction does not evolve with redshift at fixed magnitude. This lack of satellite evolution is supported broadly by observations of massive galaxies ($\mstar > 3 \times 10 ^ {10} \msun$) in X-ray-selected groups of mass $10 ^ {13 - 14} \msun$ out to $z \sim 1$ in COSMOS \citep{GeoLeaBun11}. If the satellite quiescent fraction does not evolve at fixed satellite and host halo mass, then the ratio of a satellite's quenching time to its dynamical friction lifetime must remain roughly fixed, implying that the satellite quenching time is shorter at higher redshift: $\tq \propto (1 + z) ^ {-3 / 2}$ \citep{TinWet10}. Furthermore, because the SSFR distribution of galaxies in groups is strongly bimodal out to at least $z \sim 0.4$ \citep{McGBalWil11}, this suggests that the delayed-then-rapid quenching scenario remains true at higher redshift. Also, if $\tqdelay$ decreases more quickly than $\tauqfade$, then there would be a higher fraction of satellites at intermediate SSFRs at higher redshift, a trend that is suggested by the significant fraction of `green valley' galaxies in groups at $z \sim 1$ \citep{BalMcGWil11}. In all, these result suggest that satellite quenching times are shorter at higher redshift, as we will investigate and quantify further in \citetalias{WetTinCon13b}. We also emphasize that our results are predicated on the accuracy of the SHAM method for assigning stellar mass to both satellite and central subhalos across $z = 0 - 1$. While numerous works support the (statistical) accuracy of this approach, as outlined in \S\ref{sec:galaxy_catalog_sim}, to any extent that SHAM assigns biased stellar masses to satellite subhalos at these redshifts, this would bias our derived quenching timescales. We have argued that group preprocessing is the primary reason that satellites in more massive host halos are more likely to be quiescent. Group preprocessing should manifest itself via an increased quiescent fraction for satellites that remain in a sub-group after falling into a cluster \citep{WhiCohSmi10}, to the extent that such sub-groups are observationally identifiable \citep{Coh12}. Promisingly, examining the red fraction of galaxies in clusters at $z = 0.2 - 0.5$, \citet{LiYeeEll09} saw that, within and at the outskirts of these clusters, galaxies that appear associated with groups exhibit a higher red fraction than those that are not, providing direct evidence for the importance of group preprocessing. Our results also connect with satellites of much lower mass in the Local Group. Naively extending our quenching timescales in Fig.~\ref{fig:quench-time_v_m-star} to lower mass implies that dwarf satellites in the Milky Way take \tit{at least} $5 \gyr$ to quench after infall. Given that the vast majority of satellites in the Local Group are quiescent, this implies that they fell in $> 5 \gyr$ ago, as supported by detailed comparisons of their positions and velocities to those of similar satellites in simulation \citep{RocPetBul12}. Conversely, our quenching timescales also reinforce recent claims that the Large and Small Magellanic Clouds, which are both actively star-forming, fell into the Milky Way halo within the last few Gyrs \citep{BesKalHer07}: had they fallen in much earlier, our results indicate that they would no longer be actively star-forming. Examining the stellar age distributions of satellites in the Local Group, \citet{OrbGneWei08} found that, even for those that are currently quiescent, a large fraction have experienced a significant amount of recent star formation: half of satellites with $\mstar > 10 ^ 7 \msun$ have formed more than 10\% of their stellar mass in the last $2 \gyr$. This trend, observed in lower mass satellites than we examined, supports our general result that satellites continue to form stars over extended timescales and thus grow in stellar mass considerably after infall. Examining more massive ($\mstar > 10 ^ 9 \msun$), quiescent elliptical/lenticular galaxies in the Coma cluster, \citet{TraFabDre08} found that their mean ages are identical to those of similar galaxies that are in the field, and that they have experienced star formation as recently at $z \sim 0.2$, trends that again suggest significant star formation after infall. Finally, \citet{SmiLucPri12} examined the stellar ages of quiescent satellites in the Coma cluster, finding a significant decrease in age with cluster-centric radius at low mass ($\mstar < 10 ^ {10} \msun$) but a much weaker trend at higher mass. As they argued, this trend implies that satellite-specific quenching plays a stronger role in quenching lower mass satellites, consistent with our results that more satellites were quenched as central galaxies prior to infall at higher mass. Adding a large sample of stellar ages \citep[e.g.,][]{GalChaBri05} directly to our group catalog would provide additional constraints on our derived star formation histories, as we will examine in future work.
| 12
| 6
|
1206.3571
|
1206
|
1206.2950_arXiv.txt
|
We have evaluated the contribution of the AGN population to the ionization history of the Universe based on a semi-analytic model of galaxy formation and evolution in the CDM cosmological scenario. The model connects the growth of black holes and of the ensuing AGN activity to galaxy interactions. In the model we have included a self consistent physical description of the escape of ionizing UV photons; this is based on the blast-wave model for the AGN feedback we developed in a previous paper to explain the distribution of hydrogen column densities in AGNs of various redshifts and luminosities, due to absorption by the host galaxy gas. The model predicts UV luminosity functions for AGNs which are in good agreement with those derived from the observations especially at low and intermediate redshifts ($z\sim 3$). At higher redshifts ($z>5$) the model tends to overestimate the data at faint luminosities. Critical biases both in the data and in the model are discussed to explain such apparent discrepancies. The predicted hydrogen photoionization rate as a function of redshift is found to be consistent with that derived from the observations. All that suggests to reconsider the role of the AGNs as the main driver of the ionization history of the Universe.
|
The assessment of the thermal history of the intergalactic medium at high redshifts is fundamental to understand the physical processes involved in galaxy formation and evolution, including triggering and flueling of Active Galactic Nuclei (AGNs). The ionization state of the intergalactic medium (IGM) is a function of the hydrogen ionizing UV background (UVB) conceivably produced by sources like high redshift galaxies and AGNs. Star forming galaxies are more common in the high redshift universe, and in principle can be responsible for the reionization of the IGM starting at $z>7$. However, even a first order estimate of their contribution to the ionizing UVB has two main sources of uncertainty. The first is connected with the increasing difficulty in evaluating the faint end slope of the UV luminosity function especially at the highest redshifts \citep{bouwens11,grazian11}. The second is even less understood and is connected with the poor knowledge of the average escape fraction of ionizing Lyman continuum photons from the interstellar medium of each galaxy. On the other hand, the AGN contribution is mainly affected by the poor knowledge of the faint end slope of their luminosity function especially at high redshifts due the the lack of deep AGN surveys at various wavelengths (see e.g. \citet{shankar07}). The observed escape fraction of ionizing photons in the majority of the bright AGNs is high and of the order of unity at the Lyman continuum, although the value for the faintest AGNs is unknown. Since the apparent number density of {\it bright} QSOs and AGNs is rapidly decreasing outside the redshift interval $2<z<3$ it is usually assumed that star forming galaxies can give a considerable contribution to the ionizing flux which could become dominant at $z<1$ and at $z>3$ \citep{madau91,giallo97,hama96,hama12}. In this context several attempts have been made to derive the escape fraction of UV ionizing photons both at low redshifts from space and at high redshifts $z=2-4$ from ground based observations. However, there is little evidence supporting a scenario where enough ionizing photons escape from galaxies at intermediate redshifts, and most attempts gave only upper limits on the escape fractions. Indeed recent measurements by \cite{cowie09} provided a stringent $2\sigma$ upper limit of 0.8\% for the relative escape fraction of galaxies at $0.9<z<1.4$ suggesting that galaxies are not the dominant contributors to the ionizing UV background at $z<2$. At high redshifts the situation is more difficult from the observational point of view. Direct measures of the Lyman continuum flux from galaxies at $z>4$ are increasingly difficult because of the sharp increase in the number density of intervening Lyman Limit absorptions due to optically thick clouds in the IGM. For this reason the search has been focused to the range $z\sim 3-4$, but there a discrepancy is found by different teams using different methods. Steidel and collaborators support the detection of appreciable Lyman limit flux escaping at least from the UV brightest galaxies at $z\sim 3$ \citep{steidel01}. More recent analyses suggest average escape fractions of order $10-15$\% \citep{shapley06,nestor11}, although their sample can be contaminated by foreground objects \citep{vanzella10,boutsia11}. Other teams using both spectroscopic and very deep broad or narrow band imaging gave only upper limits in the range $<5-15$\% \citep{giallo02, fernandez03,vanzella10b,boutsia11}. To summarize, the recent trends appear to favor low escape fractions from relatively bright galaxies, where the recent limits are below 1\% at low redshifts and below 10\% at $z\sim 3-4$, and cast serious doubts on any redshift evolution in the galaxy ionizing escape fraction. This leaves open the possibility that AGNs could still play a leading role in contributing to the cosmological ionizing background. The traditional view based on relatively bright QSO surveys in optical and X-ray bands lead to the standard view of an evolutionary ionizing QSO background increasing from the local value out to $z=2-3$ and then quickly decreasing (see e.g. \citet{madau99}). Previous estimates were based on simple parametric extrapolations both in luminosity and redshifts of the AGN luminosity function without any link to specific physical processes \citep{madau99}. However recent deep otpical QSO surveys at $z=3$ \citep{siana08,fontanot07} and $z=4$ \citep{glikman11} are showing the presence of a considerable number of previously unknown faint AGNs producing a rather steep luminosity function. The inclusion of X-ray detection in the selection methods enables to extend the knowledge of the luminosity function to even fainter limits \citep{fiore12} of the order of L(2-10 KeV) $> 10^{43}$ erg s$^{-1}$ out to $z> 4$. The presence of this faint population is changing our estimate of the AGN contribution to the ionizing UV background although the selection of faint AGNs at the highest redshifts $z\sim 6$ becomes difficult with the current instrumentation. Given the current observational limits, an insight into the origin of the ionizing flux at high redshifts can be gained from theoretical modelling. The latter should account for the formation and evolution of galaxies in a cosmological context, and for the co-evolution of the AGN population. Given the complex sub-grid physics involved in the AGN activity over a wide range of cosmic times and of galactic masses, semi-analytic models (SAMs) constitute a powerful tool to provide a statistically relevant sample of simulated galaxies and AGNs. Such models connect the physical processes involving gas and star formation to the merging histories of DM haloes collapsed from the primordial density field. While several SAMs including the growth of supermassive Black Holes have been presented in the literature (starting from \citet{kauffmann00}; see also \citet{croton06,bower06,marulli08,menci06,menci08}), addressing the problem of the ionizing background requires a physical description of the escape fraction in galaxies. Such a problem is deeply connected with that concerning the feedback of the AGNs onto the interstellar gas. In fact, since even a tiny amount (column densities $N_H\sim 10^{17}$ cm$^{-2}$) of galactic gas would suppress the radiation at the Lyman limit escaping from the galaxy, estimates of the escape fraction translate into computing the fraction of photons that are emitted along directions where the interstellar gas of the galaxy has been depleted. In our previous work \citep{menci08}, we have developed our SAM model (Rome-SAM, hereafter R-SAM)) to describe the expansion of the blast wave produced in the galaxy interstellar medium by AGN-driven outflows. Here we use the same model to derive the escape fraction of photons in active galaxies as the fraction of photons emitted along directions where the AGN-induced blast wave has expelled the galactic gas. When coupled with the integrated luminosity function of AGN self-consistently derived in the same model, this allows us to compute the ionizing background produced by AGNs within a cosmological model of galaxy formation. We stress that the basic quantities predicted by our computation have been extensively tested against observations in our previous works. On the one hand, the predictions concerning the growth of supermassive black holes and the corresponding evolution of the luminosity function have been compared against a wide range of different observations, from optical to X-rays, for SMBHs masses and AGN luminosities in a wide range of cosmic times extending up to z=6 for the brightest QSOs (\citet{menci03,menci04,menci06,menci08}; see also \citet{fiore12}); on the other hand, the blast-wave model for the AGN feedback allowed us to derive -within a full cosmological context and in connection with the galaxy properties- the distributions of column densities in active galaxies, as a function of redshift and of the AGN luminosity, which we successfully tested against observations \citep{menci08}. In this paper we extend the analysis of the UV AGN luminosity function to faint objects and adopt the blast-wave mechanism to predict the amount of UV photons escaping from the AGN host galaxies. The paper is organized as follows: in Sect. 2.1 we recall the basic properties of our semi-analytic model for the evolution of galaxies and AGNs. In Sect. 2.2 we describe how our blast-wave model for the AGN feedback (developed in previous papers) allows to compute the luminosity-dependent escape fraction of ionizing photons emitted by AGN. The evolution of the luminosity function of AGNs predicted by our model is compared with data in Sect. 3; the comparison with the current dataset is discussed taking into consideration possible biases in the model and in the data. In Sect. 4 we convolve the predicted intrinsic luminosity function of the AGN with the escape fraction derived in our model to derive the emissivity and the photoionization rate predicted by the model at various redshifts, and compare the latter with existing data; the contribution of AGNs with different luminosities to such observables is also shown and discussed. The final Sect. 5 is devoted to a summary and to conclusions.
|
A faint AGN population is emerging from recent multiwavelength surveys, in particular from those including X-ray detections. These are changing our perception about the AGN contribution to the ionizing UV background, although the selection of faint AGNs at the highest redshifts $z\sim 5-6$ becomes difficult with the present instrumentation. Given the current observational limits, insight into the origin of ionizing flux at high redshifts can be gained from theoretical modelling. We have adopted our semianalytic model R-SAM which successfully links the evolution of the galaxy population to that of the AGN population through the computation of the growth of supermassive black holes in the nuclei of galaxies. The model was able to reproduce several observables such as luminosity and mass functions of both galaxies and AGNs. In this context we first concentrated on the comparison between the predicted and observed LFs at the UV rest-frame wavelength $1450$ {\AA}. The agreement is satisfactory at low and intermediate redshifts up to $z\sim 3$. At higher redshifts the predicted LFs tend to overestimate the observed data by a factor $2-3$ at the fainter magnitudes $M_{1450}\sim -22$. We have discussed the possible reasons for the discrepancy. From the observational point of view it is clear that surveys of optically/UV selected AGNs which also rely on X-ray detection as a discriminant in the candidate selection process appear more complete on approaching the predicted values by the model. Indeed without the X-ray discriminant, the point-like morphological criterion is needed to avoid confusion with the more numerous Lyman break galaxy population. But this introduces significant incompleteness at faint magnitudes. Moreover several uncertainties in the selection process like e.g. the lack of spectroscopic information for the faintest AGN candidates, make the estimates of the faint end of the AGN LF appreciably uncertain at very high redshifts. Since faint AGN surveys could be affected by significant bias we have relied on our model to get an estimate of the hydrogen photoionization rate produced by the AGNs. To this aim, we have coupled our predicted luminosity functions with the physical description of the escape fraction based on the blast-wave model already introduced in our R-SAM to describe the expansion of the blast-wave produced in the interstellar medium of the hosting galaxy by the AGN-driven outflows. The blast wave clears the way for the UV photons emitted by the nuclear AGN which are free to escape outside the galaxy and ionize the intergalactic medium. To evaluate whether AGNs can ionize the intergalactic medium we have first computed the AGN emissivity at the Lyman limit as a function of luminosity and redshift. Interestingly, the predicted AGN emissivity has a definite maximum at about $M_{1450}\sim -23$ which is almost independent of redshift. Fainter sources give a negligible contribution because of the flattening of the shape of the luminosity function and because the escape fraction of UV photons decreases on average in fainter AGNs and reaches $\langle f\rangle\approx 0.1-0.3$ at $M_{1450}\sim -22$ at intermediate and high redshifts, respectively. Note that for AGNs of a given luminosity the escape fraction increases with redshift following the increase of the merging rate which is the main triggering of the QSO activity in our model. Since at $z>3$ AGNs with $M_{1450}\sim -22$ show a duty cycle of about 10\% and so represent about 10\% of the host galaxy population \citep{fiore12} then the bulk of the Lyman Continuum emitters hiding faint AGN activity represents about 1\% of the UV-dropout galaxy population similar to recent statistics on large Lyman break galaxy samples \citep{vanzella10}. The resulting photoionization rate evolves strongly from $z=0$ to $z=2$, and then remains nearly constant to a value $\Gamma_{-12} \sim 1$ up to $z\sim 4.5$. A gradual decline is apparent at higher redshifts, consistent with the trend derived from most recent analyses of the Lyman forest in QSO spectra. We have also shown that a high level of AGN ionizing emissivity up to $z\sim 6$ could still be consistent with a more gradual HeII reionization ending at $z\sim 3$, especially if absorption by HeII Lyman limit IGM clouds is taken into account. In summary, the main prediction of the model is that in a scenario where the AGN activity is linked to the dynamical galaxy evolution triggered by interactions and merging, only a mild high redshift evolution of the hydrogen photoionization rate is expected in the redshift interval $z\sim 2-5$. New deep multiwavelength surveys with accurate selection procedures are needed to build reliable AGN volume densities down to $M_{1450}\sim -23$ in the relevant redshift interval. These values are critical to understand whether AGNs can play a key role in reionizing the high redshift IGM. Our model suggests that it is time to reconsider the AGNs as the main population driving the ionization history of the Universe over the whole known redshift interval (up to $z\sim 6$) where the cosmic evolution of structures is observed.
| 12
| 6
|
1206.2950
|
1206
|
1206.2357_arXiv.txt
|
We investigate wheter there is any correlation between the X--ray afterglow luminosity and the prompt emission properties of a carefully selected sub-sample of bright {\it Swift} long Gamma--Ray Bursts (GRBs) nearly complete in redshift ($\sim 90\%$). Being free of selection effects (except flux limit), this sample provides the possibility to compare the rest frame physical properties of GRB prompt and afterglow emission in an unbiased way. The afterglow X--ray luminosities are computed at four different rest frame times (5 min, 1 hr, 11 hr and 24 hr after trigger) and compared with the prompt emission isotropic energy $E_{iso}$, the isotropic peak luminosity $L_{iso}$ and the rest frame peak energy $E_{peak}$. We find that the rest frame afterglow X--ray luminosity do correlate with these prompt emission quantities, but the significance of each correlation decreases over time. This result is in agreement with the idea that the GRB X-ray light curve can be described as the result of a combination of different components whose relative contribution and weight change with time, with the prompt and afterglow emission dominating at early and late time, respectively. In particular, we found evidence that the plateau and the shallow decay phase often observed in GRB X--ray light curves are powered by activity from the central engine. The existence of the $L_X-E_{iso}$ correlation at late times ($t_{rf} \geq 11$ hr) suggests a similar radiative efficiency among different bursts with on average about 6\% of the total kinetic energy powering the prompt emission.
|
Gamma-ray bursts (GRBs) are rapid, intense flashes of gamma-ray radiation occuring at an average rate of one event per day at cosmological distances. The high energy prompt emission is followed by a broadband (X-rays to radio ranges) afterglow (Costa et al. 1997; van Paradijs et al. 1997; Frail et al. 1997; Bremer et al. 1998; Heng et al. 2008) that can be observed up to weeks and months after the onset of the event. The {\it Swift} satellite (Gehrels et al. 2004) is operative since 2004 and provided, so far, uniform observations of prompt and afterglow emission for hundreds of GRBs. Among the most interesting uses of the {\it Swift} legacy, there are the statistical studies aimed at searching for the existence of correlations among the physical parameters of GRBs. From the study of the GRB prompt properties, robust correlations among the spectral parameters of the prompt emission and its energetic and luminosity have been found (Amati et al. 2002; Yonetoku et al. 2004; Ghirlanda, Ghisellini \& Lazzati 2004). Furthermore, many studies searching for correlations among GRB prompt and afterglow emission have been presented so far in the literature, based on the comparison of the observed properties (see, e.g. Gehrels et al. 2008 and references therein) or of the rest-frame properties, selecting GRBs for which a redshift could be measured (Berger et al. 2003; Racusin et al. 2011; Kann et al. 2010; Kann et al. 2011; Nysewander et al. 2009; Ghirlanda et al. 2009; Margutti et al. 2012; Bernardini et al. 2012a,b). Although this second approach enable to physically characterize these objects, it can be affected by observational biases, given that almost 2/3 of {\it Swift} GRBs are lacking a redshift measurement. Such correlation studies are of key importance for the understanding of the physics of GRB emission mechanisms and of their relation to their progenitors and to the surrounding environment. In this paper, we investigate the existence of correlations among afterglow emission and prompt spectral properties of a carefully selected sub-sample of {\it Swift} long GRBs presented in Salvaterra et al. (2012). This sample is nearly complete in redshift ($\sim 90\%$) and, consisting of 58 GRBs, is large enough to allow significant statistical studies. The paper is organized as follows. In section 2 we describe the properties of the GRBs composing the complete sample of Salvaterra et al. (2012). In section 3 we compare these properties and discuss our findings. Our conclusion are presented in section 4. Throughout the paper we assume a standard cosmology with $h={\Omega}_{\Lambda} = 0.7$ and ${\Omega}_{m} = 0.3$.
|
The statistical study of the rest-frame properties of GRBs gives the best opportunity to characterize the physics of these events, although such studies are often biased by the fact that almost 2/3 of GRBs are lacking a redshift measurement. In this paper, working with a sample of GRBs complete in redshift ($\sim 90\%$), we investigated the existence of correlations among the GRB X--ray afterglow luminosities and rest-frame prompt emission properties in an unbiased way. We tested the correlations between luminosities or between luminosity and energy performing a partial correlation analysis against the common dependence on the redshift (see Sect. 2.1) and obtained the following main results: - the afterglow X--ray luminosity $L_X$ at early times ($t_{rf}=5$ min and $t_{rf}=1$ hr) strongly correlate with the prompt emission isotropic energy ($E_{iso}$) and the peak luminosity ($L_{iso}$). At later times ($t_{rf}=11$ hr and $t_{rf}=24$ hr) the $L_X-E_{iso}$ and $L_X-L_{iso}$ correlations become weaker (but still significant). A similar trend is observed between $L_X$ and $E_{peak}$, even if in this case the significance of the correlation is lower at all times (although with correlation probabilities $>$ 95\%). - the strongest correlations are found comparing the early time X-ray luminosity at $t_{rf}=5$ min with $E_{iso}$ and $L_{iso}$. At this epoch the majority of the X-ray light curves display a steep or a plateau/shallow decay. This suggests that the initial steep decay and the plateau (shallow decay) phase observed in GRB X--ray light curves are powered by activity from the central engine; - GRBs showing the normal afterglow decay already at $t_{rf}=5$ min, althought expected to have high value of the Lorenz factor, do not show a significant excess in their $E_{iso}$ and $L_{iso}$ with respect to the other events; - the existence of the $L_X-E_{iso}$ correlation at late times suggests a similar radiative efficiency ($\sim$6\%) among different bursts, where brighter bursts have on average more kinetic energy; - the resulting slopes of our correlation are in agreement with the Amati and Yonetoku correlations (see also Nava et al. 2012). The inconsistency of GRB\,061021 with the $E_{peak}-E_{iso}$ correlation is likely due to the hardness of its prompt emission spectrum (high $E_{peak}$); - the decrease of significance of these correlations with time indicates that the early X--ray luminosity is still dominated by the prompt emission, while at late times the most significant contribution to the X--ray luminosity is given by the external shock afterglow emission. \begin{table*} \caption{X--ray luminosities in the 2--10 keV rest frame energy band computed at four different rest frame times (5 min, 1 hr, 11 hr and 24 hr) for the GRBs studied in this paper (see Sect. 2 for details). Errors are at the 90\% confidence level. } \centering \begin{tabular}{cccccccccc} \hline GRB & $z$ & $L_{X,5}$ & $L_{X,5}$ err & $L_{X,1}$ & $L_{X,1}$ err & $L_{X,11}$ & $L_{X,11}$ err & $L_{X,24}$ & $L_{X,24}$ err \\ & & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ & erg s$^{-1}$ \\ \hline 050318 & 1.440 & $1.966 \times 10^{48}$ & $3.932 \times 10^{47}$ & $7.197 \times 10^{46}$ & $1.583 \times 10^{46}$ & $7.054 \times 10^{44}$ & $1.622 \times 10^{44}$ & $1.202 \times 10^{44}$ & $3.245 \times 10^{43}$ \\ 050401 & 2.900 & $5.624 \times 10^{48}$ & $1.125 \times 10^{48}$ & $4.971 \times 10^{47}$ & $1.094 \times 10^{47}$ & $1.543 \times 10^{46}$ & $3.549 \times 10^{45}$ & $4.959 \times 10^{45}$ & $1.339 \times 10^{45}$ \\ 050416A & 0.650 & $2.676 \times 10^{46}$ & $5.352 \times 10^{45}$ & $4.933 \times 10^{45}$ & $1.085 \times 10^{45}$ & $6.252 \times 10^{44}$ & $1.438 \times 10^{44}$ & $2.813 \times 10^{44}$ & $7.595 \times 10^{43}$ \\ 050525A & 0.610 & $3.113 \times 10^{47}$ & $6.226 \times 10^{46}$ & $2.399 \times 10^{46}$ & $5.278 \times 10^{45}$ & $5.326 \times 10^{44}$ & $1.225 \times 10^{44}$ & $1.534 \times 10^{44}$ & $4.142 \times 10^{43}$ \\ 050922C & 2.200 & $1.725 \times 10^{48}$ & $3.450 \times 10^{47}$ & $7.660 \times 10^{46}$ & $1.685 \times 10^{46}$ & $2.880 \times 10^{45}$ & $6.624 \times 10^{44}$ & $9.855 \times 10^{44}$ & $2.661 \times 10^{44}$ \\ 060206 & 4.050 & $8.070 \times 10^{47}$ & $1.614 \times 10^{47}$ & $4.733 \times 10^{47}$ & $1.041 \times 10^{47}$ & $2.430 \times 10^{46}$ & $5.589 \times 10^{45}$ & $9.204 \times 10^{45}$ & $2.485 \times 10^{45}$ \\ 060210 & 3.910 & $1.092 \times 10^{49}$ & $2.184 \times 10^{48}$ & $1.602 \times 10^{48}$ & $3.524 \times 10^{47}$ & $7.653 \times 10^{46}$ & $1.760 \times 10^{46}$ & $2.832 \times 10^{46}$ & $7.646 \times 10^{45}$ \\ 060306 & 3.500 & $1.731 \times 10^{48}$ & $3.462 \times 10^{47}$ & $3.284 \times 10^{47}$ & $7.225 \times 10^{46}$ & $1.993 \times 10^{46}$ & $4.584 \times 10^{45}$ & $7.976 \times 10^{45}$ & $2.154 \times 10^{45}$ \\ 060614 & 0.130 & $1.883 \times 10^{46}$ & $3.766 \times 10^{45}$ & $1.928 \times 10^{44}$ & $4.242 \times 10^{43}$ & $1.214 \times 10^{44}$ & $2.792 \times 10^{43}$ & $4.114 \times 10^{43}$ & $1.111 \times 10^{43}$ \\ 060814 & 1.920 & $9.902 \times 10^{47}$ & $1.980 \times 10^{47}$ & $2.119 \times 10^{47}$ & $4.662 \times 10^{46}$ & $1.073 \times 10^{46}$ & $2.468 \times 10^{45}$ & $3.671 \times 10^{45}$ & $9.912 \times 10^{44}$ \\ 060908 & 1.880 & $9.362 \times 10^{47}$ & $1.872 \times 10^{47}$ & $2.302 \times 10^{46}$ & $5.064 \times 10^{45}$ & $6.495 \times 10^{44}$ & $1.494 \times 10^{44}$ & $2.023 \times 10^{44}$ & $5.462 \times 10^{43}$ \\ 060927 & 5.470 & $1.166 \times 10^{48}$ & $2.332 \times 10^{47}$ & $2.644 \times 10^{46}$ & $5.817 \times 10^{45}$ & $2.794 \times 10^{44}$ & $6.426 \times 10^{43}$ & $6.314 \times 10^{43}$ & $1.705 \times 10^{43}$ \\ 061007 & 1.260 & $4.946 \times 10^{48}$ & $9.892 \times 10^{47}$ & $6.865 \times 10^{46}$ & $1.510 \times 10^{46}$ & $1.116 \times 10^{45}$ & $2.567 \times 10^{44}$ & $2.904 \times 10^{44}$ & $7.841 \times 10^{43}$ \\ 061021 & 0.350 & $2.999 \times 10^{46}$ & $5.998 \times 10^{45}$ & $4.761 \times 10^{45}$ & $1.047 \times 10^{45}$ & $4.823 \times 10^{44}$ & $1.109 \times 10^{44}$ & $2.153 \times 10^{44}$ & $5.813 \times 10^{43}$ \\ 061121 & 1.310 & $1.524 \times 10^{48}$ & $3.048 \times 10^{47}$ & $2.788 \times 10^{47}$ & $6.134 \times 10^{46}$ & $1.108 \times 10^{46}$ & $2.548 \times 10^{45}$ & $3.396 \times 10^{45}$ & $9.169 \times 10^{44}$ \\ 061222A & 2.090 & $4.237 \times 10^{48}$ & $8.474 \times 10^{47}$ & $7.602 \times 10^{47}$ & $1.672 \times 10^{47}$ & $5.253 \times 10^{46}$ & $1.208 \times 10^{46}$ & $1.410 \times 10^{46}$ & $3.807 \times 10^{45}$ \\ 070521 & 1.350 & $1.512 \times 10^{48}$ & $3.024 \times 10^{47}$ & $1.068 \times 10^{47}$ & $2.350 \times 10^{46}$ & $2.207 \times 10^{45}$ & $5.076 \times 10^{44}$ & $6.208 \times 10^{44}$ & $1.676 \times 10^{44}$ \\ 071020 & 2.150 & $1.929 \times 10^{48}$ & $3.858 \times 10^{47}$ & $1.051 \times 10^{47}$ & $2.312 \times 10^{46}$ & $6.381 \times 10^{45}$ & $1.468 \times 10^{45}$ & $2.553 \times 10^{45}$ & $6.893 \times 10^{44}$ \\ 071117 & 1.330 & $2.524 \times 10^{47}$ & $5.048 \times 10^{46}$ & $2.207 \times 10^{46}$ & $4.855 \times 10^{45}$ & $2.111 \times 10^{45}$ & $4.855 \times 10^{44}$ & $9.802 \times 10^{44}$ & $2.647 \times 10^{44}$ \\ 080319B & 0.940 & $1.774 \times 10^{49}$ & $3.548 \times 10^{48}$ & $2.945 \times 10^{47}$ & $6.479 \times 10^{46}$ & $3.903 \times 10^{45}$ & $8.977 \times 10^{44}$ & $1.716 \times 10^{45}$ & $4.633 \times 10^{44}$ \\ 080319C & 1.950 & $4.485 \times 10^{48}$ & $8.970 \times 10^{47}$ & $2.291 \times 10^{47}$ & $5.040 \times 10^{46}$ & $4.512 \times 10^{45}$ & $1.038 \times 10^{45}$ & $1.250 \times 10^{45}$ & $3.375 \times 10^{44}$ \\ 080413B & 1.100 & $4.996 \times 10^{47}$ & $9.992 \times 10^{46}$ & $4.230 \times 10^{46}$ & $9.306 \times 10^{45}$ & $3.678 \times 10^{45}$ & $8.459 \times 10^{44}$ & $1.488 \times 10^{45}$ & $4.018 \times 10^{44}$ \\ 080603B & 2.690 & $2.816 \times 10^{48}$ & $5.632 \times 10^{47}$ & $1.873 \times 10^{47}$ & $4.121 \times 10^{46}$ & $1.377 \times 10^{46}$ & $3.167 \times 10^{45}$ & $5.865 \times 10^{45}$ & $1.584 \times 10^{45}$ \\ 080605 & 1.640 & $4.327 \times 10^{48}$ & $8.654 \times 10^{47}$ & $2.052 \times 10^{47}$ & $4.514 \times 10^{46}$ & $3.258 \times 10^{45}$ & $7.493 \times 10^{44}$ & $8.409 \times 10^{44}$ & $2.270 \times 10^{44}$ \\ 080607 & 3.040 & $8.135 \times 10^{48}$ & $1.627 \times 10^{48}$ & $1.560 \times 10^{47}$ & $3.432 \times 10^{46}$ & $3.464 \times 10^{45}$ & $7.967 \times 10^{44}$ & $9.975 \times 10^{44}$ & $2.693 \times 10^{44}$ \\ 080721 & 2.590 & $3.240 \times 10^{49}$ & $6.480 \times 10^{48}$ & $1.320 \times 10^{48}$ & $2.904 \times 10^{47}$ & $2.664 \times 10^{46}$ & $6.127 \times 10^{45}$ & $7.436 \times 10^{45}$ & $2.008 \times 10^{45}$ \\ 080804 & 2.200 & $7.173 \times 10^{47}$ & $1.435 \times 10^{47}$ & $4.539 \times 10^{46}$ & $9.986 \times 10^{45}$ & $3.181 \times 10^{45}$ & $7.316 \times 10^{44}$ & $1.334 \times 10^{45}$ & $3.602 \times 10^{44}$ \\ 080916A & 0.690 & $2.784 \times 10^{46}$ & $5.568 \times 10^{45}$ & $7.932 \times 10^{45}$ & $1.745 \times 10^{45}$ & $6.859 \times 10^{44}$ & $1.578 \times 10^{44}$ & $3.015 \times 10^{44}$ & $8.141 \times 10^{43}$ \\ 081007 & 0.530 & $3.102 \times 10^{46}$ & $6.204 \times 10^{45}$ & $4.460 \times 10^{45}$ & $9.812 \times 10^{44}$ & $6.888 \times 10^{44}$ & $1.584 \times 10^{44}$ & $3.050 \times 10^{44}$ & $8.235 \times 10^{43}$ \\ 081121 & 2.510 & $2.135 \times 10^{49}$ & $4.270 \times 10^{48}$ & $5.657 \times 10^{47}$ & $1.245 \times 10^{47}$ & $1.715 \times 10^{46}$ & $3.945 \times 10^{45}$ & $5.468 \times 10^{45}$ & $1.476 \times 10^{45}$ \\ 081203A & 2.100 & $3.081 \times 10^{48}$ & $6.162 \times 10^{47}$ & $1.508 \times 10^{47}$ & $3.318 \times 10^{46}$ & $1.519 \times 10^{45}$ & $3.494 \times 10^{44}$ & $3.379 \times 10^{44}$ & $9.123 \times 10^{43}$ \\ 081221 & 2.260 & $1.167 \times 10^{49}$ & $2.334 \times 10^{48}$ & $3.630 \times 10^{47}$ & $7.986 \times 10^{46}$ & $2.416 \times 10^{46}$ & $5.557 \times 10^{45}$ & $1.013 \times 10^{46}$ & $2.735 \times 10^{45}$ \\ 081222 & 2.770 & $4.728 \times 10^{48}$ & $9.456 \times 10^{47}$ & $2.776 \times 10^{47}$ & $6.107 \times 10^{46}$ & $1.186 \times 10^{46}$ & $2.728 \times 10^{45}$ & $2.498 \times 10^{45}$ & $6.745 \times 10^{44}$ \\ 090102 & 1.550 & $4.146 \times 10^{48}$ & $8.292 \times 10^{47}$ & $1.270 \times 10^{47}$ & $2.794 \times 10^{46}$ & $3.943 \times 10^{45}$ & $9.069 \times 10^{44}$ & $1.267 \times 10^{45}$ & $3.421 \times 10^{44}$ \\ 090424 & 0.540 & $5.712 \times 10^{47}$ & $1.142 \times 10^{47}$ & $6.465 \times 10^{46}$ & $1.422 \times 10^{46}$ & $3.566 \times 10^{45}$ & $8.202 \times 10^{44}$ & $1.383 \times 10^{45}$ & $3.734 \times 10^{44}$ \\ 090715B & 3.000 & $3.764 \times 10^{48}$ & $7.528 \times 10^{47}$ & $1.724 \times 10^{47}$ & $3.793 \times 10^{46}$ & $1.403 \times 10^{46}$ & $3.227 \times 10^{45}$ & $4.369 \times 10^{45}$ & $1.180 \times 10^{45}$ \\ 090812 & 2.450 & $2.409 \times 10^{48}$ & $4.818 \times 10^{47}$ & $1.735 \times 10^{47}$ & $3.817 \times 10^{46}$ & $6.525 \times 10^{45}$ & $1.501 \times 10^{45}$ & $2.232 \times 10^{45}$ & $6.026 \times 10^{44}$ \\ 090926B & 1.240 & $4.025 \times 10^{47}$ & $8.050 \times 10^{46}$ & $1.991 \times 10^{46}$ & $4.380 \times 10^{45}$ & $1.125 \times 10^{45}$ & $2.588 \times 10^{44}$ & $4.396 \times 10^{44}$ & $1.187 \times 10^{44}$ \\ 091018 & 0.970 & $8.393 \times 10^{47}$ & $1.679 \times 10^{47}$ & $3.749 \times 10^{46}$ & $8.248 \times 10^{45}$ & $1.879 \times 10^{45}$ & $4.322 \times 10^{44}$ & $7.062 \times 10^{44}$ & $1.907 \times 10^{44}$ \\ 091020 & 1.710 & $1.419 \times 10^{48}$ & $2.838 \times 10^{47}$ & $1.151 \times 10^{47}$ & $2.532 \times 10^{46}$ & $4.329 \times 10^{45}$ & $9.957 \times 10^{44}$ & $1.481 \times 10^{45}$ & $3.999 \times 10^{44}$ \\ 091127 & 0.490 & $9.707 \times 10^{47}$ & $1.941 \times 10^{47}$ & $8.700 \times 10^{46}$ & $1.914 \times 10^{46}$ & $5.367 \times 10^{45}$ & $1.234 \times 10^{45}$ & $1.646 \times 10^{45}$ & $4.444 \times 10^{44}$ \\ 091208B & 1.060 & $3.825 \times 10^{47}$ & $7.650 \times 10^{46}$ & $3.422 \times 10^{46}$ & $7.528 \times 10^{45}$ & $1.980 \times 10^{45}$ & $4.554 \times 10^{44}$ & $7.799 \times 10^{44}$ & $2.106 \times 10^{44}$ \\ 100621A & 0.540 & $2.438 \times 10^{47}$ & $4.876 \times 10^{46}$ & $5.484 \times 10^{46}$ & $1.206 \times 10^{46}$ & $4.382 \times 10^{45}$ & $1.008 \times 10^{45}$ & $1.558 \times 10^{45}$ & $4.207 \times 10^{44}$ \\ 100728B & 2.106 & $2.891 \times 10^{47}$ & $5.782 \times 10^{46}$ & $3.204 \times 10^{46}$ & $7.049 \times 10^{45}$ & $8.411 \times 10^{44}$ & $1.935 \times 10^{44}$ & $2.559 \times 10^{44}$ & $6.909 \times 10^{43}$ \\ 110205A & 2.220 & $7.380 \times 10^{48}$ & $1.476 \times 10^{48}$ & $1.160 \times 10^{47}$ & $2.552 \times 10^{46}$ & $2.126 \times 10^{45}$ & $4.890 \times 10^{44}$ & $5.752 \times 10^{44}$ & $1.553 \times 10^{44}$ \\ 110503A & 1.613 & $2.480 \times 10^{48}$ & $4.960 \times 10^{47}$ & $1.733 \times 10^{47}$ & $3.813 \times 10^{46}$ & $7.646 \times 10^{45}$ & $1.759 \times 10^{45}$ & $2.516 \times 10^{45}$ & $6.793 \times 10^{44}$ \\ \hline \hline \end{tabular} \label{tab_log1} \end{table*} \begin{table*} \centering \caption{Correlation fits and coefficients. Data distributions were fitted with the function $y = 10^Ax^B$ (see Sect. 2.1 for details). Errors are at $1\sigma$ confidence level.} \begin{tabular}{cccccc} \hline Correlation & A & B & $r$ & $P_{null}$ & Dispersion \\ & & & & & \\ \hline $L_{X,5}$ $vs.$ $E_{iso}$ & $-6.31 \pm 2.70$ & $1.03 \pm 0.04$ & 0.80 & $1.44 \times 10^{-12}$ & 0.256 \\ $L_{X,1}$ $vs.$ $E_{iso}$ & $-3.30 \pm 3.69$ & $0.95 \pm 0.05$ & 0.63 & $1.65 \times 10^{-6}$ & 0.341 \\ $L_{X,11}$ $vs.$ $E_{iso}$ & $ 1.51 \pm 5.55$ & $0.83 \pm 0.07$ & 0.36 & $1.45 \times 10^{-2}$ & 0.436 \\ $L_{X,24}$ $vs.$ $E_{iso}$ & $-0.21 \pm 6.92$ & $0.86 \pm 0.08$ & 0.31 & $3.37 \times 10^{-2}$ & 0.477 \\ $L_{X,5}$ $vs.$ $L_{iso}$ & $ 1.44 \pm 3.06$ & $0.89 \pm 0.04$ & 0.59 & $1.06 \times 10^{-6}$ & 0.337 \\ $L_{X,1}$ $vs.$ $L_{iso}$ & $ 3.85 \pm 3.04$ & $0.82 \pm 0.04$ & 0.51 & $2.47 \times 10^{-4}$ & 0.335 \\ $L_{X,11}$ $vs.$ $L_{iso}$ & $ 7.70 \pm 4.35$ & $0.72 \pm 0.05$ & 0.36 & $1.38 \times 10^{-2}$ & 0.428 \\ $L_{X,24}$ $vs.$ $L_{iso}$ & $ 6.06 \pm 4.98$ & $0.74 \pm 0.06$ & 0.35 & $1.93 \times 10^{-2}$ & 0.467 \\ $L_{X,5}$ $vs.$ $E_{peak}$ & $43.84 \pm 6.74$ & $1.62 \pm 0.16$ & 0.49 & $5.98 \times 10^{-4}$ & 0.351 \\ $L_{X,1}$ $vs.$ $E_{peak}$ & $43.01 \pm 6.74$ & $1.49 \pm 0.15$ & 0.36 & $1.40 \times 10^{-2}$ & 0.358 \\ $L_{X,11}$ $vs.$ $E_{peak}$ & $42.31 \pm 7.46$ & $1.23 \pm 0.16$ & 0.15 & $3.31 \times 10^{-1}$ & 0.407 \\ $L_{X,24}$ $vs.$ $E_{peak}$ & $41.84 \pm 7.53$ & $1.23 \pm 0.17$ & 0.11 & $4.71 \times 10^{-1}$ & 0.427 \\ \hline \hline \end{tabular} \label{tab:log_corr} \end{table*} \begin{table*} \centering \caption{Correlation fits and coefficients for the GRBs of our sample divided into two redshift bins ($z < 1.8$ and $z> 1.8 $; see Sec. 3 for details). Data distributions were fitted with the function $y = 10^Ax^B$ (see Sect. 2.1 for details). Errors are at $1\sigma$ confidence level.} \begin{tabular}{cccccc} \hline Correlation & A & B & $r$ & $P_{null}$ & Dispersion \\ & & & & & \\ \hline \multicolumn{6}{|c|}{$z < 1.8$} \\ \hline $L_{X,5}$ $vs.$ $E_{iso}$ & $-5.14 \pm 3.97$ & $1.01 \pm 0.05$ & 0.68 & $2.72 \times 10^{-4}$ & 0.284 \\ $L_{X,1}$ $vs.$ $E_{iso}$ & $-0.45 \pm 4.98$ & $0.89 \pm 0.06$ & 0.73 & $5.92 \times 10^{-5}$ & 0.351 \\ $L_{X,11}$ $vs.$ $E_{iso}$ & $13.91 \pm 6.12$ & $0.60 \pm 0.06$ & 0.51 & $1.38 \times 10^{-2}$ & 0.355 \\ $L_{X,24}$ $vs.$ $E_{iso}$ & $12.13 \pm 7.25$ & $0.62 \pm 0.07$ & 0.49 & $1.91 \times 10^{-2}$ & 0.399 \\ $L_{X,5}$ $vs.$ $L_{iso}$ & $ 1.84 \pm 4.26$ & $0.88 \pm 0.05$ & 0.66 & $5.25 \times 10^{-4}$ & 0.350 \\ $L_{X,1}$ $vs.$ $L_{iso}$ & $ 5.88 \pm 3.41$ & $0.78 \pm 0.04$ & 0.70 & $1.84 \times 10^{-4}$ & 0.283 \\ $L_{X,11}$ $vs.$ $L_{iso}$ & $18.89 \pm 4.54$ & $0.51 \pm 0.04$ & 0.53 & $9.84 \times 10^{-3}$ & 0.290 \\ $L_{X,24}$ $vs.$ $L_{iso}$ & $17.66 \pm 5.27$ & $0.52 \pm 0.04$ & 0.58 & $4.21 \times 10^{-3}$ & 0.334 \\ $L_{X,5}$ $vs.$ $E_{peak}$ & $44.08 \pm 10.29$ & $1.50 \pm 0.23$ & 0.37 & $9.09 \times 10^{-2}$ & 0.417 \\ $L_{X,1}$ $vs.$ $E_{peak}$ & $43.30 \pm 9.86$ & $1.35 \pm 0.20$ & 0.37 & $8.88 \times 10^{-2}$ & 0.403 \\ $L_{X,11}$ $vs.$ $E_{peak}$ & $43.02 \pm 11.07$ & $0.93 \pm 0.15$ & 0.22 & $3.39 \times 10^{-1}$ & 0.364 \\ $L_{X,24}$ $vs.$ $E_{peak}$ & $42.54 \pm 11.27$ & $0.94 \pm 0.16$ & 0.30 & $1.80 \times 10^{-1}$ & 0.374 \\ \hline \multicolumn{6}{|c|}{$z > 1.8$} \\ \hline $L_{X,5}$ $vs.$ $E_{iso}$ & $ -7.77 \pm 6.18$ & $1.06 \pm 0.08$ & 0.77 & $9.76 \times 10^{-6}$ & 0.230 \\ $L_{X,1}$ $vs.$ $E_{iso}$ & $-8.43 \pm 12.55$ & $1.05 \pm 0.08$ & 0.44 & $3.82 \times 10^{-2}$ & 0.333 \\ $L_{X,11}$ $vs.$ $E_{iso}$ & $-17.12 \pm 30.94$ & $1.18 \pm 0.24$ & 0.28 & $2.18 \times 10^{-1}$ & 0.438 \\ $L_{X,24}$ $vs.$ $E_{iso}$ & $-18.31 \pm 41.03$ & $1.19 \pm 0.27$ & 0.19 & $3.99 \times 10^{-1}$ & 0.472 \\ $L_{X,5}$ $vs.$ $L_{iso}$ & $ -3.54 \pm 10.00$ & $0.98 \pm 0.13$ & 0.48 & $2.35 \times 10^{-2}$ & 0.329 \\ $L_{X,1}$ $vs.$ $L_{iso}$ & $ -4.82 \pm 16.80$ & $0.98 \pm 0.17$ & 0.30 & $1.78 \times 10^{-1}$ & 0.383 \\ $L_{X,11}$ $vs.$ $L_{iso}$ & $-13.60 \pm 30.76$ & $1.12 \pm 0.23$ & 0.26 & $2.43 \times 10^{-1}$ & 0.461 \\ $L_{X,24}$ $vs.$ $L_{iso}$ & $-15.72 \pm 36.14$ & $1.15 \pm 0.25$ & 0.22 & $3.20 \times 10^{-1}$ & 0.485 \\ $L_{X,5}$ $vs.$ $E_{peak}$ & $44.34 \pm 12.53$ & $1.46 \pm 0.36$ & 0.38 & $7.78 \times 10^{-2}$ & 0.299 \\ $L_{X,1}$ $vs.$ $E_{peak}$ & $43.71 \pm 10.27$ & $1.26 \pm 0.39$ & 0.19 & $3.94 \times 10^{-1}$ & 0.342 \\ $L_{X,11}$ $vs.$ $E_{peak}$ & $43.08 \pm 65.28$ & $0.97 \pm 0.53$ & 0.06 & $8.06 \times 10^{-1}$ & 0.484 \\ $L_{X,24}$ $vs.$ $E_{peak}$ & $42.87 \pm 21.20$ & $0.87 \pm 0.56$ &-0.06 & $8.01 \times 10^{-1}$ & 0.537 \\ \hline \hline \end{tabular} \label{tab:log_corr} \end{table*}
| 12
| 6
|
1206.2357
|
1206
|
1206.6604_arXiv.txt
|
We investigate the instability of purely poloidal magnetic fields in nonrotating neutron stars by means of three-dimensional general-relativistic magnetohydrodynamics simulations, extending the work presented in \citet{Ciolfi2011}. Our aim is to draw a clear picture of the dynamics associated with the instability and to study the final configuration reached by the system, thus obtaining indications on possible equilibria in a magnetized neutron star. Furthermore, since the internal rearrangement of magnetic fields is a highly dynamical process, which has been suggested to be behind magnetar giant flares, our simulations can provide a realistic estimate of the electromagnetic and gravitational-wave emission which should accompany the flare event. Our main findings are the following: (i) the initial development of the instability meets all the expectations of perturbative studies in terms of the location of the seed of the instability, the timescale for its growth and the generation of a toroidal component; (ii) in the subsequent nonlinear reorganization of the system, $\sim 90$\% of magnetic energy is lost in few Alfv\'en timescales mainly through electromagnetic emission, and further decreases on a much longer timescale; (iii) all stellar models tend to achieve a significant amount of magnetic helicity and the equipartition of energy between poloidal and toroidal magnetic fields, and evolve to a new configuration which does not show a subsequent instability on dynamical or Alfv\'en timescales; (iv) the electromagnetic emission matches the duration of the initial burst in luminosity observed in giant flares, giving support to the internal rearrangement scenario; (v) only a small fraction of the energy released during the process is converted into $f$-mode oscillations and in the consequent gravitational-wave emission, thus resulting in very low chances of detecting this signal with present and near future ground based detectors.
|
Neutron stars (NSs) are endowed with very intense, long-lived, large-scale magnetic fields, reaching strengths which are estimated to be of the order of $10^{13}$ G at the magnetic pole for ordinary NSs, and around $10^{15}$ G in the case of magnetars. Such extreme magnetic fields play a crucial role in the physics of NSs, affecting their structure and evolution. They are involved in the processes through which NSs are observed, like the pulsar magnetic dipole radiation and the magnetically-powered burst activity of magnetars, and they have been recently recognised as essential to explain the quasi-periodic oscillations detected in the aftermath of magnetar giant flares [see, \eg, \citet{Gabler2012} and references therein]. Moreover, they are responsible for deformations which may cause a significant emission of gravitational waves \citep{Bonazzola1996,Cutler2002} and precession \citep{Wasserman2003}, they influence the thermal evolution of the star \citep{Pons2009}, to list a few. All these processes depend on the magnetic field configuration inside the NSs, whose geometry is basically unknown. From observations of the spindown in pulsars the exterior magnetic field appears to be purely poloidal and mainly dipolar, but substantially different internal geometries can reproduce this external appearance. The importance of obtaining such information has motivated a significant effort in studying possible equilibrium models of magnetized NSs, at first with simple geometries, \eg purely poloidal or purely toroidal fields, and recently with mixed poloidal-toroidal fields. The latest models, built in Newtonian and general-relativistic framework, include \citet{Tomimura2005, Lander:2009, Lander2012, Ciolfi2009, Ciolfi2010, Fuji2012}, where the so-called `twisted-torus' configuration is considered. This particular geometry has been found as a result of the evolution of initial random fields in Newtonian magnetohydrodynamic (MHD) simulations by \citet{BraithNord2006}. Once a magnetic-field geometry is chosen, building a corresponding equilibrium configuration is not sufficient to assess whether this represents a good description of the NS interior. The magnetic field, in fact, should also be long-lived and thus stable on timescales which are much longer than the dynamical timescale. Assessing the stability of a given magnetic field configuration is not trivial and most of the work done on the subject concerns simple field geometries and nonrotating stars. Until very recently, the problem has been only addressed with a perturbative analytic approach. These calculations established that both a purely poloidal field and a purely toroidal field are unstable in nonrotating stars, giving important predictions about the onset of the instability, but they could not predict the following evolution of the system. Only recently, thanks to the progress in numerical simulations, it has become possible to study these hydromagnetic instabilities by performing the fully three-dimensional (3D) MHD evolutions of magnetized relativistic stars. These simulations represent a very powerful tool, allowing to confirm the predicted features of the instability and to obtain information about the nonlinear dynamics of the process. In addition, the end-state of simulations can provide important hints about the preferred magnetic field configuration in magnetized stars. There is an additional and important motivation for studying hydromagnetic instabilities in NSs. The induced global rearrangement of magnetic fields is a violent, strongly dynamical process, and soon after the magnetar model was proposed \citep{DT92}, this kind of process was suggested as a trigger mechanism of giant flares \citep{TD95,TD2001}. Nowadays, this internal rearrangement scenario represents one of the leading models to explain the phenomenology observed in magnetars, the other one involving a large-scale rearrangement of magnetic fields in the magnetosphere surrounding the star~\citep{Lyutikov2003,Lyutikov2006,Gill2010}. Since both mechanisms can be present in a giant flare, the main question becomes whether most of the magnetic energy powering the flare is stored inside the star or in its exterior magnetosphere. The basic tests on these models rely on the comparison of the predicted timescales and energies involved with the giant flare observations. Hence, determining self-consistently the dynamics associated to this kind of instability can provide important hints on the underlying mechanism. Moreover, magnetar flares (and in particular giant flares) are likely to be accompanied by a significant excitation of NS oscillations, in particular in the $f$-mode, which can then lead to a strong emission of gravitational waves (GWs). This possibility has motivated recent searches for GWs in connection to magnetar flares, published by the LIGO and Virgo collaboration [see, \eg \citet{ligo-virgo2011}]. Semi-analytic efforts have been devoted to establishing the maximum amount of magnetic energy released in a magnetar flare, which, in turn, provides an upper limit on the energy emitted in GWs \citep{Ioka01,Corsi2011}. These upper limits are based on analytical calculations and simplified models, and can only provide rough, order-of-magnitude estimates. The assumption that all the available energy (which is at most of the order of the total magnetic energy) is converted into GWs, leads to the optimistic conclusion that the signal would be detectable with the next-generation ground-based detectors \citep{Corsi2011}. This result has been questioned in \citet{Levin:2011}, where a simple perturbative analysis is employed to show that only a small fraction of the magnetic energy involved in a giant flare event can be actually be converted into $f$-mode oscillations and that the consequent GW emission is expected to be very weak. Again, referring to the internal magnetic field rearrangement scenario of giant flares, MHD simulations of hydromagnetic instabilities can provide a realistic picture of the GW signal produced, together with estimates of the fraction of energy which can be pumped into the $f$-mode and of the signal detectability. In this work we focus on the instability of purely poloidal fields, with the goal of shedding some light on all of the points made above\footnote{The instability of purely-toroidal fields has several analogies with the one considered here for purely-poloidal fields and has been investigated by \citet{Kiuchi:2008,Kiuchi:2011}.}. Two parallel works back in the '70s~\citep{Markey1973,Wright1973} found that poloidal fields suffer from the so-called ``Tayler'' or ``kink'' instability, which manifests itself firstly in the neighbourhood of the neutral line. This instability was recently studied with Newtonian numerical simulations in the linear regime \citep{Lander:2011}, or with nonlinear evolutions for a simplified model of newly born NS \citep{Geppert2006}, and in the case of main-sequence stars \citep{Braithwaite2007}. The first 3D general-relativistic MHD simulations of the poloidal field instability in NSs were presented only last year, in two parallel works \citep{Lasky2011,Ciolfi2011}. These studies reported similar results, despite some substantial difference in the approach (in particular in the evolution of magnetic fields outside the star), essentially confirming all of the analytic predictions on the instability, and providing some first hints about the nonlinear rearrangement of the system. In addition, \citet{Ciolfi2011} presented the first examples of gravitational waveforms triggered by the instability, which were subsequently considered more systematically by \citet{Zink:2011} for nonrotating stars and by \citet{Lasky2012} for rotating NSs. In this paper we extend the work presented in \citet{Ciolfi2011}, presenting additional information on the numerical infrastructure used and considering the instability-driven evolution of a series of nonrotating NSs endowed with purely poloidal magnetic fields of different strength. The organization of the paper is as follows. In Sect.~\ref{setup} we reconsider the setup of the system, improving in particular our treatment of the atmosphere. Within the new setup, we confirm that our evolutions meet all the expectations on the onset of the instability, in agreement with the previous perturbative studies. In Sect.~\ref{evolution} we examine in more detail the nonlinear rearrangement of the system, performing much longer simulations and gaining new substantial insight on the final state reached by the system. Section~\ref{endstate} is dedicated to a discussion of the implications for the most-likely magnetic-field configurations in magnetized NSs and of the role played by magnetic helicity. We also study the emission properties of the system, relevant for the internal field rearrangement scenario of giant flares, estimating in Sect.~\ref{ememission} the timescale of the process and its electromagnetic luminosity, and discussing the detectability of the GW signal in Sect.~\ref{gravitationalwaves}. Both for electromagnetic and GW emissions, our estimates rely on a good agreement with the dependence on the magnetic field strength expected in the perturbative limit of weak magnetic fields, which allows us to extrapolate our results to lower and more realistic values than those actually considered in the simulations. Our conclusions are finally presented in Sect.~\ref{sec:conclusions}. Unless specified differently, we adopt units in which $c=1$, $G=1$.
|
\label{sec:conclusions} We have performed 3D general-relativistic MHD simulations of nonrotating magnetized NSs endowed with a purely poloidal magnetic field and studied the development of the hydromagnetic instability that develops dynamically. This work represents an extension of our previous study on the subject \citep{Ciolfi2011}, where we have improved our treatment of the atmosphere, drastically reducing the undesired energy losses due to the transition between the ideally conducting stellar interior and the resistive exterior, and we have performed simulations on much longer timescales, which have allowed us to gain essential information about the final configuration reached by the system. The resulting overall picture is much clearer and conclusive. As expected from perturbation analyses, the instability is first triggered in the region of close field lines and is accompanied by the production of a toroidal magnetic field. When the growth of the latter saturates in about one Alfv\'en timescale, the toroidal field has reached a local strength which is comparable to the poloidal one and the axisymmetry of the initial configuration is lost. At this point, the most dynamical phase of the nonlinear evolution takes place, with major modifications of the magnetic field leading to a strong electromagnetic emission carrying away $\sim 90\%$ of the magnetic energy in few Alfv\'en timescales. At the same time, a small fraction of magnetic energy is converted into stellar oscillations, mostly at the $f$-mode frequency, which cause the emission of gravitational waves. The subsequent evolution proceeds on longer timescales, with further loss of magnetic energy. Since in the post-instability phase the magnetic field is continuously changing, losing strength because of resistive dissipation, it is difficult to determine whether the corresponding configuration is a stationary one or not. The only robust evidence is that the variations in the hydrodynamical and electromagnetic quantities are much smaller than those during the instability, taking place on larger timescales. We therefore interpret this behaviour as evidence that the new magnetic-field configuration has reached a quasi-stationary state or that, if still intrinsically unstable, the growth time of the instability is much larger than the one that can be possibly investigated numerically. Because in this quasi-stationary configuration the ratio of the toroidal and poloidal magnetic energies tends to unity, and because the growth of toroidal magnetic field at the expense of the poloidal one is also accompanied by an increase in the magnetic helicity, we are led to conclude that an equilibrium magnetic field configuration with a significant amount of magnetic helicity and comparable poloidal and toroidal magnetic field energies could be a preferred one for stability. Bearing this in mind, we remark that it is still unclear if stable equilibria exist for a simple barotropic fluid star, and that other stabilizing contributions, such as a stable stratification, may be necessary to obtain long-lived magnetic field configurations. The violent reorganization of magnetic fields induced by the development of a hydromagnetic instability has been proposed as a possible mechanism to explain giant flares in magnetars. Using our simulations and in particular the information about the dissipated magnetic energy, we have deduced, within the approximation of our resistive approach, an estimate of the electromagnetic luminosity $L_{em}$ associated with the hydromagnetic instability. More specifically, we found that the average luminosity scales as $\propto B^3$ with the magnetic-field strength, in good agreement with the expectation that the radiated energy should scale as $\propto B^2$, while the duration of the emission should scale as $\propto B^{-1}$. In this way we were able to perform a direct comparison with the observations of SGR 1806-20, finding a very good agreement with the duration of the burst and an emitted luminosity which is about an order of magnitude larger than the measured one. As a result, although our modelling is oversimplified, the overall agreement with the observations from SGR 1806-20, lends support to the suggestion that the basic phenomenology is that of an internal-field readjustment. Finally, we have discussed the gravitational-wave emission which should be expected as a result of the $f$-mode oscillations triggered by the instability. Also in this case, our calculations reproduce the expected scaling behaviour of the root-mean-square gravitational-wave amplitude in the limit of weak magnetic fields, \ie $h_{\rm rss} \sim B^2$. This important result allows us to extrapolate with confidence our estimates to smaller magnetic fields than those covered by our simulations and conclude that the gravitational-wave signal from $f$-mode oscillations will be undetectable for realistic values of the magnetic field, \ie $B_{\rm p} \lesssim 10^{15}\,{\rm G}$ and marginally detectable by third generation detectors for unrealistically large magnetic fields, \ie $B_{\rm p} \gtrsim 2\times 10^{16}\,{\rm G}$. As a self-consistent solution to this problem in fully resistive regime is close to be within reach~\citep{Kiki2012}, we plan to extend the investigation reported above and thus remove many of the approximations that our present analysis had to sustain. It will then be possible to set even more precise connections between the complex phenomenology observed in magnetar flares and the violent dynamics of hydromagnetic instabilities in relativistic stars. \bigskip We are grateful to D. Radice, A. Harte, S.~K. Lander, S. Mereghetti, W. Kastaun, B. Giacomazzo, E. Bentivegna, G.~M. Manca, R. De Pietri and S. Bernuzzi for useful discussions. R. Ciolfi is supported by the Humboldt Foundation. Support comes also from ``CompStar'', a Research Networking Programme of the European Science Foundation and by the DFG grant SFB/Transregio~7. The calculations have been performed on the supercomputing clusters at the AEI.
| 12
| 6
|
1206.6604
|
1206
|
1206.4437_arXiv.txt
|
s{ I consider the prospects for performing weak lensing studies with the new generation of radio telescopes that are coming online now and in the future. I include a description of a proposed technique to use polarization observations in radio weak lensing analyses which could prove extremely useful for removing a contaminating signal from intrinsic alignments. Ultimately, the Square Kilometre Array promises to be an exceptional instrument for performing weak lensing studies due to the high resolution, large area surveys which it will perform. In the nearer term, the e-MERLIN instrument in the UK offers the high sensitivity and sub-arcsec resolution required to prove weak lensing techniques in the radio band. I describe the SuperCLASS survey -- a recently accepted e-MERLIN legacy programme which will perform a pioneering radio weak lensing analysis of a supercluster of galaxies.}
|
I have given a brief summary of the status of the field of weak lensing in the radio band. While it currently lags well behind the field of optical weak lensing, the new radio instruments coming online now make radio weak lensing a viable alternative which is complementary to large scale optical surveys. In particular, radio polarization observations offer interesting possibilities for removing intrinsic alignments from radio lensing surveys. Over the course of the next few years, the SuperCLASS survey on the e-MERLIN telescope will act as a pathfinder experiment for more ambitious radio lensing surveys with future instruments.
| 12
| 6
|
1206.4437
|
|
1206
|
1206.6118_arXiv.txt
|
The newly discovered galaxy cluster 1RXS J0603.3+4214 hosts a 1.9 Mpc long, bright radio relic with a peculiar linear morphology. Using hydrodynamical N-body AMR simulations of the merger between three initially hydrostatic clusters in an idealised setup, we are able to reconstruct the morphology of the radio relic. Based on our simulation, we can constrain the merger geometry, predict lensing mass measurements and X-ray observations. Comparing such models to X-ray, redshift and lensing data will validate the geometry of this complex merger which helps to constrain the parameters for shock acceleration of electrons that produces the radio relic.
|
Some merging galaxy clusters host diffuse extended radio emission, so-called radio halos and relics, unrelated to individual galaxies. The origin of these halos and relics is still debated, although there is compelling evidence that they are related to cluster mergers \citep{ensslin98, miniati01, hoeft07, hoeft08, pfrommer08, battaglia09, skillman11, vazza12}. Van Weeren (2012) present detailed Westerbork Synthesis Radio Telescope (WSRT) and Giant Metrewave Radio Telescope (GMRT) radio observations between 147 MHz and 4.9 GHz of a new radio-selected galaxy cluster 1RXS J0603.3+4214, located at a redshift of 0.225. The cluster hosts a large bright 1.9 Mpc radio relic, so-called toothbrush for its shape, as well as an elongated $\sim 2$ Mpc radio halo, and two fainter radio relics (van Weeren et al. 2012 submitted). Part of the main radio relic has a peculiar linear morphology. The cluster is detected as an extended X-ray source in the ROSAT All Sky Survey with an X-ray luminosity of $L_{X, 0.1-2.4 {\rm keV}}\sim 1\times10^{45}$ erg s$^{-1}$. An overlay of the radio and X-ray data is shown in Fig.~\ref{fig:tooth}. The distorted intracluster medium (ICM) is clearly elongated along the N-S and possibly in the E-W directions, indicating a complex merger. V,R, and I band 4.2m William Herschel Telescope (WHT) images confirmed the presence of a large galaxy cluster. For the bright radio relic a clear spectral index gradient from the front of the relic towards the cluster centre is observed. Parts of the relic are highly polarized with a polarization fraction of up to 60\%. A model in which particles are (re)accelerated in a first-order Fermi process at the front of the relic provides the best match to the observed spectra. The orientation of the bright relic and halo indicate a north-south merger, but the peculiar linear shape and the presence of another relic, perpendicular to the bright relic, suggest a more complex merger. Deep X-ray observations as well as lensing studies will be needed to completely understand the dynamics of this cluster. Here we attempt to model this peculiar object using the hydrodynamic AMR code FLASH v3.3. In particular, we would like to model (i) the linear morphology of the relic and (ii) its orientation with respect to the X-ray emission. As a result, we will be making predictions about the mass profile that may be detected via gravitational lensing. Understanding the dynamics of the gas in this object is crucial for understanding the physical conditions under which relics are formed. There exist a number of simulations of cluster mergers \citep{1993ApJ...407L..53R,1993A&A...272..137S, 1994ApJ...427L..87B, 1994MNRAS.268..953P, 1997ApJS..109..307R, 1998ApJ...496..670R,1999ApJ...518..594R,2000ApJ...538...92R, 2001ApJ...561..621R, 2000ApJ...535..586T, 2002MNRAS.329..675R, 2006MNRAS.373..881P,2009ApJ...699.1004Z, 2010ApJ...717..908Z, 2011ApJ...728...54Z, vanweeren11}. Cluster mergers can decouple the baryonic matter component from the dark matter (DM) which causes an offset between the gravitational centre (measured from lensing) and X-ray centre of the cluster. This was first observed for the ``Bullet cluster'' \citep[1E0657$-$56,][]{2006ApJ...648L.109C}. \cite{2007MNRAS.380..911S} presented hydrodynamical models of galaxy cluster mergers to reproduce the dynamical state and mass models (from gravitational lensing) of the ``Bullet'' cluster (1E0657$-$56). \cite{2008MNRAS.389..967M} presented detailed N-body/SPH simulations of the system. In the following, we describe our attempts to reproduce the event that could form the "toothbrush cluster". \vspace{-0.5cm}
| 12
| 6
|
1206.6118
|
|
1206
|
1206.6742_arXiv.txt
|
Massive binary stars may constitute a substantial fraction of progenitors to supernovae and $\gamma$-ray bursts, and the distribution of their orbital characteristics holds clues to the formation process of massive stars. As a contribution to securing statistics on OB-type binaries, we report the discovery and orbital parameters for five new systems as part of the Cygnus~OB2 Radial Velocity Survey. Four of the new systems (MT070, MT174, MT267, and MT734$\equiv$VI Cygni \#11) are single-lined spectroscopic binaries while one (MT103) is a double-lined system (B1V+B2V). MT070 is noteworthy as the longest period system yet measured in Cyg~OB2, with $P$=6.2~yr. The other four systems have periods ranging between 4 and 73~days. MT174 is noteworthy for having a probable mass ratio $q<0.1$, making it a candidate progenitor to a low-mass X-ray binary. These measurements bring the total number of massive binaries in Cyg~OB2 to 25, the most currently known in any single cluster or association.
|
Massive stars produce some of nature's most energetic phenomena as they end their lives in spectacular supernova or $\gamma$-ray burst explosions \citep{burrows95,woosleylangerweaver,woosley2006}. While the post-main-sequence evolutionary path of a single massive star is still not settled (e.g., the sequence through red/blue supergiant, Wolf-Rayet, Luminous Blue Variable phases for stars of various masses; \citealt[][]{meynet2000}), the possible evolutionary paths of close massive binary systems are even less certain. Nevertheless, close massive binaries appear common \citep{garmany80,sanaevans2011,Kiminki2012a}, and secondary stars in such systems have been implicated in removing hydrogen envelopes from massive primaries prior to core collapse, resulting in a substantial fraction (30--40\%) of Type Ib/c supernovae and $\gamma$-ray bursts \citep{izzard2004, kf07, eldridge2008}. Close massive binaries may also produce the observed population of high-velocity ``runaway'' stars either through ejection of one component after the supernova explosion or during multi-body gravitational interactions \citep{giesbolton, blaauw}. Papers I--IV in this series (\citealt{Kiminki07}, \citealt{Kiminki08}, \citealt{Kiminki09}, \citealt{Kiminki2012a}) describe prior results of the Cygnus~OB2 Radial Velocity Survey intended to measure the massive binary characteristics (i.e., binary fraction, distribution of periods, mass ratios, eccentricities) for a large number (114) massive stars in a single cluster/association having a common formation environment. In particular, Paper IV (Table~6) summarizes orbital elements for the 20 massive binaries with measured parameters. Paper V \citep{Kiminki2012b} uses these previously published data and the new data presented herein (a total of 25 measured systems) to infer the intrinsic distributions of massive binaries, concluding that the fraction of massive stars (defined as B3 and earlier; i.e., supernova progenitors\footnote{For a typical initial mass function \citep{salpeter55}, B stars outnumber O stars 3:1 as supernova progenitors.}) having companions may be as high as 90\%, and that 45\% of these are likely to interact at some point. Paper V reports an excess of short-period 4--7 day systems relative to 7--14 day systems and finds that the distribution of mass ratios is approximately flat over the range 0.1$<q<$1.0. Besides providing the fundamental data for modeling the frequencies of energetic phenomena, these types of statistics help place constraints on theoretical frameworks for the formation of massive stars which remain under debate \citep{Krumholzetal2010, smithetal2009}. In this sixth paper of the series we present orbital solutions for five additional massive spectroscopic binaries in Cyg~OB2---one double-lined (SB2) and four single-lined (SB1) systems---using spectroscopic data obtained primarily during 2010--2011, but utilizing some data as early as 1999. Using the nomenclature of \citet{MT91}, these systems are MT103 (SB2), MT070, MT174, MT267, and MT734 (SB1s). Figure~\ref{color} displays a three-color representation of the Cyg~OB2 vicinity, where blue/green/red depict the Palomar Sky Survey R, the $Spitzer$ 4.5~$\mu$m, and the $Spitzer$ 8.0~$\mu$m bands, respectively. White labels denote previously known binary systems, while magenta labels highlight the newly discovered systems reported herein. Numeration follows the system of \citet{MT91}, with ``S'' additionally indicating the numeration of \citet{Schulte58} and ``A'' or ``B'' for that of \citet{comeron02}. Figure~\ref{color} shows the complex nature of this region, including ionized gas (diffuse blue emission tracing H$\alpha$ within the POSS R band), photo-dissociation regions (PDR) at the edges of molecular clouds (diffuse red and green tracing broad emission features arising from excited polycyclic aromatic hydrocarbons, PAHs), and stars (blue and green point sources). After first reviewing the Survey strategy and data, we present the new radial velocity data and describe the orbital characteristics of these O and early B systems. We conclude by summarizing the parameters of the 25 currently known massive binaries in CygOB2. All reported velocities are in the heliocentric frame unless explicitly indicated otherwise.
|
We have presented orbital solutions for five additional massive binaries that are probable members of the Cygnus~OB2 Association, bringing the total number of multiple systems therein to 25. The vast majority of these have complete orbital solutions, making this the largest collection of massive binary statistics in any single cluster or association to date. \citet{Kiminki2012b} have used these data, along with previously known systems documented in Papers I--IV, to infer the underlying distribution of periods, mass ratios, and eccentricities for Cyg~OB2 as a whole. The analysis of \citet{Kiminki2012b} suggests that the known list of 25 binaries out of 114 Cyg~OB2 systems surveyed is reasonably complete at periods less than about 30~days, but that we should expect to discover an additional 10--15 binaries having longer periods as the Cygnus~OB2 Radial Velocity Survey continues. Noteworthy among the new list of massive binaries is the discovery of a probable progenitor of a low-mass X-ray binary system (MT174). Also noteworthy is the long-period system MT070 with a period of 6.2 years. Both of these low-amplitude systems ($K_1$$\simeq$9 \kms) attest to the capabilities of the Survey to achieve orbital solutions for either low-inclination or low-$q$ systems.
| 12
| 6
|
1206.6742
|
1206
|
1206.1481_arXiv.txt
|
The measurement of polarization of the background stars in the region of Bok globules is important to study the magnetic field geometry and dust grain characteristics in the globule. These parameters are important for the formation and evolution of dark clouds. We made polarimetric observations of Bok globule CB56 in the R-filter from the 2-metre telescope at IUCAA Girawali Observatory (IGO). The observations were carried out on 2011 March 4th and 5th. The CCD images obtained from the instrument (IFOSC) were analyzed, to produce the polarization map of the Bok globule CB56.
|
\label{s:intro} Bok globules are the most simple molecular clouds in our Milky Way Galaxy, which are ideal sites for low-mass star formation (Bok \& Reilly 1947). These are small, opaque, and relatively isolated molecular clouds with diameters of about 0.7 pc (0.1 -- 2 pc) and masses of $\approx 10$ M$_\odot$ (2 -- 100 M$_\odot$) (Bok 1977, and Leung 1985). Several catalogues of globules and dark clouds were published during the past years (e.g. Barnard 1927; Bok 1956; Lynds 1962; Sandquist \& Lindroos 1976; Feitzinger \& Stuewe 1984; Hartly et al. 1986; Clemens \& Barvainis 1988; Persi et al. 1990). The most homogeneous and complete compilation of dark clouds has been done by Clemens \& Barvainis (1988) (henceforth CB catalogue). This catalogue contains 248 small (mean size $\sim 4^\prime$) and nearby (distance $<$ 1 kpc) molecular clouds. Magnetic fields play a major role on the evolution of dark clouds and may control the fragmentation of clouds to form stars (Mestel \& Spitzer 1956; Nakano \& Nakamura 1978; Mouschovias \& Morton 1991; Li \& Nakamura 2004). Polarization arises from alignment of interstellar dust grains in the presence of a magnetic field (Davis \& Greenstein 1951). When the light from the background stars passes through the clouds, extinction and reddening are caused due to absorption and scattering by the dust grains present in the clouds. This phenomenon also introduces linear polarization in the starlight, if the dust grains are aligned and dichroic. Observations of the polarization of background stars define the morphology of the magnetic field in the plane of the sky, and provide deep insights into the effect of the field on the geometrical structure of a collapsing dark cloud (Goodman et al. 1990). Polarimetric studies of these clouds provide important information about the optical properties of the dust grains as a function of environment. There have been many studies of polarization with a view to trace the geometry of the magnetic fields in dark clouds. Bhatt \& Jain (1993) presented linear polarization measurements of stars in the regions of the molecular clouds B227 and L121. Bhatt \& Jain (1993) also discussed the magnetic field geometry in the clouds as indicated by the polarization maps in relation to their morphology. Kane et al. (1995) studied Bok globule CB4 using a CCD imaging polarimeter in order to create a detailed map of the magnetic field associated with CB4 cloud. Sen et al. (2000) have mapped eight star-forming clouds, CB3, CB25, CB39, CB52, CB54, CB58, CB62 and CB246, in white light polarization and commented on the possible star formation dynamics there. Sen et al. (2005) modelled the dark cloud as a simple dichroic polarizing sphere, which explains why polarization need not always increase with total extinction A$_V$ as one moves towards the center of the cloud. Their analysis shows that the observed polarization depends largely on the orientation of the magnetic field (within the cloud) with respect to the direction of interstellar magnetic field. Ward-Thompson et al. (2009) studied optical and submillimetre polarimetric data of the Bok globule CB3 and CB246. They found that the field orientation deduced from the optical polarization data matches well with the orientation estimated from the submillimetre measurements of the Bok globule CB3. Recently, Sen et al. (2010) studied three clouds CB3, CB25 and CB39 photopolarimetrically to find any possible relation between the observed polarization and extinction of the stars in the background of these clouds. They found that the measured extinction values increase with the increase in percentage polarization for the cloud CB39 and to some extent for CB25. However, they did not observe any such correlation for cloud CB3. In this paper, we present the results of polarization measurements for stars in the region of the Bok globule CB56. This cloud is compact, irregular shaped and has two infrared point sources in the IRAS Point Source Catalogue (IRAS 07125-2503 and 07125-2507) associated with it (Clemens \& Barvainis 1988). CB56 is an ideal site for star formation. Since no optical polarimetric observation of CB56 has yet been made, we select this globule as our target object. It is to be noted that the polarimetric observations of dark cloud in the optical range help us to infer the magnetic field orientation in the low density edge regions of clouds, whereas the submillimetre polarimetric observation of dark cloud can only trace the field orientation in the high density central regions of the clouds.
|
We have presented the results of linear polarization measurement for stars in the region of the dark cloud CB56. The polarization map of this cloud indicates that the magnetic field in CB56 is more or less unidirectional and nearly parallel to the long axis of the cloud, and has the same direction as the local interstellar magnetic field (Clemens \& Barvainis 1988).
| 12
| 6
|
1206.1481
|
1206
|
1206.6097_arXiv.txt
|
{Corona, Chromosphere, Magnetic Fields, Coronal Heating} Determining the heating mechanism (or mechanisms) that causes the outer atmosphere of the Sun, and many other stars, to reach temperatures orders of magnitude higher than their surface temperatures has long been a key problem. For decades the problem has been known as the coronal heating problem, but it is now clear that `coronal heating' cannot be treated or explained in isolation and that the heating of the whole solar atmosphere must be studied as a highly coupled system. The magnetic field of the star is known to play a key role, but, despite significant advancements in solar telescopes, computing power and much greater understanding of theoretical mechanisms, the question of which mechanism or mechanisms are the dominant supplier of energy to the chromosphere and corona is still open. Following substantial recent progress, we consider the most likely contenders and discuss the key factors that have made, and still make, determining the actual (coronal) heating mechanism (or mechanisms) so difficult.
|
The coronal heating problem is one of the longest running solar physics puzzles and is still a highly controversial topic. Considerable progress has been made in modelling possible heating mechanisms, but discriminating amongst these to discover if any one process dominates is extremely difficult to do. This is because the corona is not energetically isolated from the other regions of the atmosphere, such as the chromosphere, but instead the whole of the solar atmosphere forms a highly coupled system, with both energy and mass transferred in both directions between the chromosphere and corona through the transition region. The chromosphere has very different plasma properties to the corona and so explaining the heating of the solar atmosphere is now recognised to be considerably more complicated than had been appreciated during much of the twentieth century. The coronal heating problem first arose following the work of \citet{Grotrian1939} (erroneously the author on the paper is given as W. Grotian) and \citet{Edlen1942} who discovered that emission lines observed during a total solar eclipse in 1869 were not due to a new element called coronium, but were the result of highly ionised iron. This work demonstrated that the temperature of the corona is in excess of a million degrees Kelvin. In comparison, the temperature near the Sun's surface, the photosphere, is just 6,000 K. At the same time the density falls off by six orders of magnitude from the photosphere to the corona. The Sun is not unique in having such a hot outer atmosphere. Indeed, most stars in the cool half of the Hertzsprung-Russell (H-R) diagram have a corona and can be observed in X-rays. Their average temperatures range from 1-45 million degrees Kelvin with the hottest associated with the most rapidly rotating stars. Using observations from the Einstein Observatory, \citet{vaiana1981} was the first to show that most stars, as opposed to just the odd exotic star, have soft X-ray emission and, hence, a corona. \citet{linsky1985} reviewed stars across the H-R diagram with emissions in ultraviolet and X-ray emission lines which are produced by plasma hotter than $10^4$ K. He found that Dwarf stars of spectral type G-M and rapidly rotating subgiants and giants of spectral type F-K in spectroscopic binary systems are most likely to be solar-like in nature, i.e., they are likely to have ``a turbulent magnetic field sufficiently strong to control the dynamics and energetics of their outer atmosphere''. Furthermore, he found that T Tauri stars, other pre-Main-Sequence stars and single giants of spectral type F to early K are also probably solar-like. \begin{figure}[ht] \centering{ \includegraphics[scale=0.112]{parnellf1a} \includegraphics[scale=0.112]{parnellf1b} \includegraphics[scale=0.112]{parnellf1c}} \centering{ \includegraphics[scale=0.112]{parnellf1d} \includegraphics[scale=0.112]{parnellf1e} \includegraphics[scale=0.112]{parnellf1f} } \caption{(a-e) SDO/AIA 211 \AA, 193 \AA, 094 \AA, 335 \AA\ and 304 \AA\ images of the solar corona and chromosphere showing the amazing complexity and vast range of scales of structures. (f) SDO/HMI magnetogram of the magnetic field in the photosphere. All images taken on 27th August 2011. \label{fig:sdosun}} \end{figure} X-ray, EUV and UV images of the Sun (Figure~\ref{fig:sdosun}(a-e)) show beautiful, finely detailed loops structures, as well as more diffuse emission which is likely to be formed from a mass of unresolved loops. The resolved loops either connect regions of opposite-polarity magnetic field seen in the photosphere, or they are extended open structures that are anchored in a photospheric magnetic feature. As a general rule, the hotter the loops, the stronger the regions of magnetic field they are associated with \citep{Fisher1998}. The structure of the magnetic field is highlighted in coronal images because heat is spread efficiently along field lines, but transport across them is greatly inhibited (thermal conduction in the corona is 13 orders of magnitude stronger along field lines than across them \citep{Priest1982}). All of these factors signify that the magnetic field plays a major role in heating the solar corona. Moreover, there is strong evidence from other stars that X-ray coronae are associated with magnetic field since a power-law relation exists between X-ray luminosity and magnetic flux over many orders of magnitude, ranging from solar quiet-regions, through active regions, to dwarf and TTauri stars \citep{Fisher1998,Pevtsov2003}. The loops observed, for example in Figure~\ref{fig:sdosun}(a-e), are visible since they contain plasma of the appropriate density and temperature to emit at the given wavelength. The surrounding space contains loops that are either at a different temperature (hotter or cooler) or insufficiently dense to be observed. In equilibrium, and without any heating, a loop will be cool and rarefied, so any heating mechanism must explain both how the material fills the loop and heats it to coronal temperatures. In the corona, the combined radiative and conductive losses in active regions are $10^7$ erg cm$^{-2}$ s$^{-1}$ and in the quiet-Sun are $3\times10^5$ erg cm$^{-2}$ s$^{-1}$, whereas in the chromosphere they are $2\times 10^7$ erg cm$^{-2}$ s$^{-1}$ and $4\times 10^6$ erg cm$^{-2}$ s$^{-1}$ in active regions and the quiet-Sun, respectively \citep{Withbroe1977}. The source of this energy is convective churning of the plasma at and below the photosphere. These motions continuously move the footpoints of the ubiquitous magnetic fields that thread through the surface of the Sun with the timescale of these motions, relative to the end-to-end Alfv{\'e}n travel time, leading to one of two fundamental types of heating mechanism. Slow (long timescale) motions result in a quasi-static stressing of the field, whilst fast (short timescale) motions generate waves. The dissipation of magnetic stresses, which typically manifest themselves in the form of current sheets, leads to magnetic reconnection and is known as DC heating, whilst the dissipation of waves is referred to as AC heating. In both cases the actual dissipation will occur at kinetic scales, since the Lundquist number (the ratio of the Ohmic dissipation time to the Alfv{\'e}n crossing time) is very large ($\approx 10^{10-13}$). On the other-hand, the photospheric driving and the complex geometry/topology of the magnetic field are determined globally. Not surprisingly, this massive range of scales which need to be addressed by coronal heating models poses a problem for both theorists and observers. Current parallel computers have neither the memory nor the power to permit well resolved models on all scales and current imagining telescopes have neither the required spatial and temporal resolution nor the sensitivity to see anything at the kinetic scale. Additional difficulties arise in the form of the coupling between the dense interior and tenuous outer atmosphere which require very different physics and, for open magnetic field regions, mechanisms must also account for the mass flux loss due to the solar wind. With all these difficulties it is not surprising that, to date, researchers have been unable to properly explain the solar heating problem. Nonetheless our understanding has advanced significantly. Since the identification of the hot solar atmosphere, there have been a huge number of heating mechanisms proposed that involve reconnection or magnetic waves. A full review of the analytical and numerical studies carried out is far beyond the scope of this review, so we refer the interested reader to the reviews by \citet{Aschwanden2004}, \citet{Goedbloed2004}, \cite{Klimchuk2006}, \citet{Aschwanden2007}, \citet{Goossens2011} and \citet{Reale2010}. Instead, we focus on those mechanisms currently most favoured. First, though, in Section 2, we consider the source of the corona's energy, before giving a brief selective discussion on the back grounds to wave heating (Section 3) and magnetic reconnection (Section 4). Nanoflare observations are reviewed in Section 5 and the heating of coronal loops, including numerical modelling and the recent observational results on spicules and Alfv\'enic waves, is the focus of Section 6. Finally, in Section 7, we present our conclusions.
|
For decades after the discovery of the very high temperatures in the solar atmosphere, solving the coronal heating problem was mostly framed as an `AC or DC question', or `waves versus reconnection'. Within the last few years, observational results have shown that the coronal heating problem has to be thought of in a different way. The highly dynamical nature of the solar atmosphere will clearly lead to the occurrence of both types of processes and the true question is to determine the relative contribution of the different mechanisms in different coronal structures. Wave-based heating mechanisms very much seem to have come full circle, from being the first mechanism suggested to provide the necessary energy input into the solar atmosphere, to being mostly discarded and recently back to the forefront again, driven by the spectacular observations by Hinode and SDO. Additionally, the discovery of the abundance of Type II spicules demands that the impact of the equally dynamic chromosphere has to be assessed much more carefully and that it is essential that the `coronal' heating question must be studied as part of the highly coupled solar atmosphere. Indeed, it is well known (but often ignored) that the energy requirements of the chromosphere far exceed those of the corona. Heating by multiple small-scale reconnection events has long been considered a viable mechanism and much support has been lent to this mechanism both observationally and theoretically, but definitive observations of the smoking-gun are still missing. Indeed for all coronal heating mechanisms there are still issues that need to be addressed before a definitive conclusion can be reached. For instance, the inability to directly `observe' coronal heating has meant that proxies have had to be found. However, it is very difficult using these proxies to distinguish absolutely between wave heating or reconnection mechanisms. For example, the drifting of heating layers in phase mixing/resonant absorption means that such a wave-based mechanism will actually be impulsive in nature and so appear very much like a nanoflare heating event \citep{Ofman1998,Moriyasu2004,Klimchuk2006,Mendoza2006}. Recent studies by Antolin and co-authors (\citet{Antolin2008,Antolinetal2010}) investigate the different observational signatures of coronal heating by either nanoflares or the dissipation of Alfv\'en waves. These authors suggest that coronal rain could be a marker for coronal heating mechanisms, shedding light not only on the spatial distribution of heating events in coronal loops but potentially also on the actual heating mechanism itself. Additionally, on the theoretical front it is clear that simplified models are not sufficient for comparison with observations. The response of the plasma in the solar atmosphere is very complex and interpreting responses of emission lines is not simple either. Forward modelling of theoretical experiments which have physically realistic effects included is essential, but the analysis of these results must be done with considerable care to determine which are important effects and will take time to complete. In the meantime, the argument over which mechanism(s) heat(s) the corona will continue.
| 12
| 6
|
1206.6097
|
1206
|
1206.4571_arXiv.txt
|
Infrared 3.6 to 8~\micron\ images of the Extended Groth Strip yield plausible counterpart identifications for all but one of 510 radio sources in the AEGIS20 $S(1.4~{\rm GHz})>50$~\mmjy\ sample. This is the first such deep sample that has been effectively 100\% identified. Achieving the same identification rate at $R$-band would require observations reaching $R_{AB}>27$. Spectroscopic redshifts are available for 46\% of the sample and photometric redshifts for an additional 47\%. Almost all of the sources with 3.6~\micron\ AB magnitudes brighter than 19 have spectroscopic redshifts $z<1.1$, while fainter objects predominantly have photometric redshifts with $1\la z\la3$. Unlike more powerful radio sources that are hosted by galaxies having large stellar masses within a relatively narrow range, the AEGIS20 counterparts have stellar masses spanning more than a factor of 10 at $z\sim1$. The sources are roughly 10--15\% starbursts at $z\la0.5$ and 20--25\% AGNs mostly at $z>1$ with the remainder of uncertain nature.
|
Radio observations are an excellent way to identify star-forming galaxies and active galactic nuclei (AGNs). Radio surveys are not subject to selection effects of obscuration or spectral line contamination, which affect visible-light surveys. Even very distant objects can have large radio flux densities. However, radio surveys alone are not sufficient to understand the populations, and followup observations are often more difficult than the initial radio survey. The problem is that redshift $z>1$ sources are faint in visible light and require very deep followup studies in order to achieve identifications. Despite the difficulties of counterpart identification, there are now several radio samples with identification rates of $\ga$90\%. \citet{Waddington2000} found optical counterparts for 96\% of the sources in a $S(1.4~{\rm GHz})>1$~mJy sample with images reaching $R=26$. \citet{Ciliegi2003} found counterparts for 92\% of $S(6~{\rm cm})>50$~\mmjy\ sources with $I_{\rm AB}=25$ images. \citet{Afonso2006} were able to identify only 89\% of an $S(1.4~{\rm GHz})>61$~\mmjy\ radio sample even with \hst/ACS observations reaching magnitude $z_{850}=28$. \citet{Simpson2006} identified $>$90\% of a 100~\mmjy\ radio sample using $BRi'z'$ images reaching AB magnitude 27. \citet{Mainieri2008}, using data over a wide wavelength range including the infrared, chose counterparts for 95\% of a 42--125~\mmjy\ radio sample but with an estimated 3\% rate of spurious identifications. In other recent work, \citet{Bardelli2010} reported an 82\% identification rate in a $\sim$50~\mmjy\ radio sample in the COSMOS field, and \citet{Afonso2011} achieved 83\% identification of a sample of ultra-steep-spectrum radio sources with $S(610~{\rm MHz})>100$~\mmjy\ in the Lockman Hole. \citet{Huynh2008} identified 79\% of a much fainter ($S(1.4~{\rm GHz})>10$~\mmjy) radio sample using deep HST images ($I_{\rm AB}<26$) albeit with a relatively large matching radius (up to 1\farcs96). Even when counterparts are detected, observed visible light corresponds to rest-frame ultraviolet for high redshift galaxies. While this can give a measure of star formation {\em rate}, it gives little indication of stellar {\em mass} and thus little indication of the type of galaxy hosting the radio source. Radio-quiet AGNs and star-forming galaxies both contribute to the faint radio population, but lack of complete identifications and limited wavelength coverage make the proportions uncertain \citep[e.g.,][]{Huynh2008}. The All-wavelength Extended Groth strip International Survey (AEGIS) \citep{Davis2007} offers an unprecedented combination of deep, multiwavelength data over a large area, the Extended Groth Strip (EGS). The data include a radio survey at 20~cm, AEGIS20 \citep{Ivison2006}, which reaches a sensitivity limit of 50~\mmjy~beam$^{-1}$. \citet{Willner2006} showed that infrared observations of radio sources can produce very high identification rates for radio source counterparts albeit at much higher flux densities (55~mJy~beam$^{-1}$) than AEGIS20 and at 6~cm rather than 20~cm. \citet{Park2008} claimed to find all the AEGIS20 radio sources that are also 24~\micron\ sources on the IRAC images but used a 2\farcs5 matching radius. IRAC data at 20~minute depth contributed to the \citet{Afonso2011} identifications in the Lockman Hole. Infrared data should give high identification rates because the SED of typical stellar populations peaks near 1.6~\micron. For a distant source, this peak will be redshifted to longer wavelength, and the IRAC 3.6 and 4.5~\micron\ flux densities will not decrease as rapidly as might be expected. Passive evolution also increases the observed flux densities in these band: a stellar population of a given mass was brighter in the past when it was younger. (See Fig.~1 of \citealt{Eisenhardt2008}.) Thus counterparts should be visible in IRAC observations unless they are either extremely distant, have very low mass, or are heavily obscured by dust, neither of the latter two being likely for a powerful radio galaxy. This paper reports the matching of AEGIS20 radio sources primarily to IRAC 3.6 to 8.0~\micron\ data. \citet{Barmby2008} have provided images and catalogs of IRAC data in the EGS. The typical exposure time is 2.5~hours (9~ks), and the 80\% completeness limits for the catalog are $\sim$5~\mmjy\ at 3.6 and 4.5~\micron\ and $\sim$10~\mmjy\ at 5.8 and 8.0~\micron. However, fainter objects with known positions can be identified on the images. Additional IRAC data for the EGS exist (Ashby et al., in preparation, 2012) but were not used for this work. In practice, the \citeauthor{Barmby2008} data suffice to identify counterparts for all or nearly all of the radio sources. The radio sample is defined and source matching is described in Section~\ref{s:id}, counterpart properties including photometry and redshifts are given in Section~\ref{s:prop}, and results are discussed in Section~\ref{s:disc} and summarized in Section~\ref{s:conc}. Throughout this paper, magnitudes are in the AB system, and the notation $[w]$ means the AB magnitude at wavelength $w$ in \micron.\footnote{Some other papers use the $[w]$ notation to mean Vega magnitudes, but here it means AB.} Source distances are based on standard $\Lambda$CDM cosmology with $H_0=71$~km~s$^{-1}$~Mpc and $\Omega_M = 0.27$. Practical calculation of luminosity distances was based on the program {\sc angsiz} \citep{Kayser1997}.
|
\label{s:conc} IRAC images are a powerful means of identifying and classifying radio sources. Images with $\sim$9~ks IRAC depth giving $\sigma\approx0.1$~\mmjy\ \citep{Barmby2008} are sufficient to detect essentially all counterparts of radio sources in a sample with 1.4~GHz brightness limit of 50~\mmjy/beam. Radio sources at this depth are roughly 10--15\% local ($z\la0.5$) starbursts and 20--25\% AGNs mostly at $z>1$ with the remainder of uncertain nature. More than 1/3 of the sample have counterparts with $R_{\rm AB}>24$, and 15\% have $R_{\rm AB}>25.5$. These sources would be very difficult to identify in $R$-band surveys, and if simple position matching were used, many would be incorrectly identified with brighter objects that are nearby on the sky but unrelated to the radio source. The AEGIS20 sample, now essentially 100\% identified, offers a great opportunity for detailed studies of the radio population. X-ray observations and additional spectra should enable a better classification between star-forming and active-nuclei galaxies, and the sample should yield luminosity functions and evolutionary history of the various populations.
| 12
| 6
|
1206.4571
|
1206
|
1206.3484_arXiv.txt
|
We study planetesimal evolution in circumbinary disks, focusing on the three systems Kepler 16, 34 and 35 where planets have been discovered recently. We show that for circumbinary planetesimals, in addition to secular forcing, eccentricities evolve on a dynamical timescale, which leads to orbital crossings even in the presence of gas drag. This makes the current locations of the circumbinary Kepler planets hostile to planetesimal accretion. We then present results from simulations including planetesimal formation and dust accretion, and show that even in the most favourable case of 100\% efficient dust accretion, in situ growth starting from planetesimals smaller than $\sim{10}~\mathrm{km}$ is difficult for Kepler 16b, Kepler 34b and Kepler 35b. These planets were likely assembled further out in the disk, and migrated inward to their current location.
|
In the last decade planets have been found in very perturbed systems such as close binary star systems. The first of these planets to be discovered were orbiting the primary star \citep{queloz00,hatzes03,zucker04}, but the latest additions to the family, after promising results using stellar eclipse timings \citep{lee09}, involve planets in circumbinary orbits: Kepler 16 \citep{doyle11} and Kepler 34 and 35 \citep{welsh12}. The parameters of these new planets are summarised in Table \ref{tab1}. The existence of planets in these systems baffles planet formation theory. A crucial step in the process of building a planet, namely growing gravitationally bound protoplanets from km-sized planetesimals, can be hindered or stopped in these perturbed environments for planetesimals on circumprimary orbits \citep{marzari00,thebault06,paard08,thebault11}. The coupling between gravitational perturbations of the companion star and gas drag stirs up the eccentricities of planetesimals, which leads to high encounter velocities. This makes accretion towards larger bodies difficult. Similar problems haunt planetesimals on circumbinary orbits\citep{moriwaki04,scholl07,marzari08,meschiari12}. Above studies focused on gravitational dynamics and gas drag only. In this work, we investigate the effect of collisions on the evolution of the system. \cite{collision} showed that in a system with high-speed collisions, it is necessary to keep track of collision outcomes. Notably, if collisions are mostly destructive, any surviving planetesimals are embedded in a sea of small debris. If they can pick up some of this debris, planetesimals can grow despite the hostile environment \citep{collision,xie10}. In this Letter, we aim to explore this possibility in the newly found planet-harbouring systems of Kepler 16, 34 and 35. We begin in Section \ref{secEcc} by reviewing the eccentricity evolution of planetesimals in circumbinary orbits. We discuss the model in Section \ref{secMod}, present the results in Section \ref{secRes}, and conclude in Section \ref{secDisc}. \begin{deluxetable}{rrrr} \tablecolumns{4} \tablewidth{0pc} \tablecaption{Binary and planet parameters} \tablehead{ \colhead{ } & \colhead{Kepler 16$^a$} & \colhead{Kepler 34$^b$} & \colhead{Kepler 35$^b$} } \startdata $M_A/\msun$ & 0.69 & 1.0 & 0.89\\ $M_B/\msun$ & 0.20 & 1.0 & 0.81\\ $\ab/$AU & 0.22 & 0.23 & 0.18\\ $\eb$ & 0.16 & 0.52 & 0.14\\ $M_\mathrm{p}/M_\mathrm{J}$ & 0.33 & 0.22 & 0.13\\ $\aP/\ab$ & 3.2 & 4.7 & 3.4\\ $\ep$ & 0.0069 & 0.18 & 0.042 \enddata \tablenotetext{a}{\cite{doyle11}} \tablenotetext{b}{\cite{welsh12}} \label{tab1} \end{deluxetable}
|
\label{secDisc} We have studied planetesimal collisions in circumbinary gas disks, focusing on the planet-harbouring systems Kepler 16, 34 and 35. We have shown that in addition to secular forcing, planetesimals experience eccentricity forcing on a dynamical timescale, which leads to eccentricity oscillations and orbital crossings that can not be prevented by gas drag. This makes the current location of the planets Kepler 16b, 34b and 35b very hostile for planetesimal accretion. We then used a numerical model similar to that of \cite{collision} including planetesimal formation and accretion of small dust. Even in the most favourable case of 100\% efficient dust accretion, we have been unable to grow planetesimals from initially 1 km at the current location of the planets. Since dust accretion is likely to be less than 100\% efficient, for example because not all the small dust will be concentrated in the midplane of the disk, we conclude that in situ planetesimal accretion is difficult for the planets Kepler 16b, 34b and 35b. We have made several necessary simplifications to make following the collisional evolution of the planetesimal population tractable. First of all, we have ignored gas dynamics throughout and worked with a static circular gas disk. While the gas disk is likely to become eccentric, especially at small radii, it was shown in \cite{paard08} that unless the gas relaxes towards the forced eccentricity, including gas dynamics makes matters worse for planetesimal accretion. For the fast eccentricity oscillations to be damped by gas drag, the gas disk will have to oscillate in phase with the planetesimals. Full hydrodynamical simulations are necessary to determine whether this is the case. These can also be used to study the effect of the inner truncation of the gas disk, and we will consider such simulations in future investigations. We considered the planar case, but a small inclination of the binary plane with respect to the gas disk may promote planetesimal accretion in the circumprimary case \citep{xie09}. However, because of the fast eccentricity oscillations and the resulting orbital crossings it is unclear if this effect can play a role in the circumbinary case. Moreover, it was shown in \cite{fragner11} that including gas dynamics again makes matters worse, even in the misaligned disk case. A formation mechanism which can leapfrog the problematic km-size range, such as gravitational collapse aided by streaming instabilities \citep{johansen07}, may overcome the problems of planetesimal accretion. It remains to be seen, however, if such a mechanism can operate in close binary systems. Preliminary calculations show that in the current model, we would need to start with planetesimals of at least $10~\mathrm{km}$ in order for in situ accretion of the Kepler circumbinary planets to become possible. The most straightforward solution is that the three circumbinary planets were assembled further out in an accretion-friendly region, and migrated in towards their current location at a later stage. This can be achieved at a relatively early stage, in the 10-100 km size range, by radial drift due to a pressure gradient in the gas, or at a later stage when the planet is more or less fully grown, by Type I or Type II planetary migration. Whatever the migration mechanism, it is likely that the inner edge of the truncated gas disk will cause migration to stall. We then expect the current location of the planets to be close to the truncation radius of the gas disk.
| 12
| 6
|
1206.3484
|
1206
|
1206.4565_arXiv.txt
|
{We present a new search for variable stars in the Galactic globular cluster M28 (NGC~6626).} {The search is based on a series of $BVI$ images obtained with the SMARTS Consortium's 1.3m telescope at Cerro Tololo Inter-American Observatory, Chile.} {The search was carried out using the ISIS v2.2 image subtraction package. } {We find a total of 25 variable stars in the field of the cluster, 9 being new discoveries. Of the newly found variables, 1 is an ab-type RR Lyrae star, 6 are c-type RR Lyrae, and 2 are long-period/semi-regular variables. V22, previously classified as a type II Cepheid, appears as a bona-fide RRc in our data. In turn, V20, previously classified as an ab-type RR Lyrae, could not be properly phased with any reasonable period.} {The properties of the ab-type RR Lyrae stars in M28 appear most consistent with an Oosterhoff-intermediate classification, which is unusual for bona-fide Galactic globulars clusters. However, the cluster's c-type variables do not clearly support such an Oosterhoff type, and a hybrid Oosterhoff I/II system is accordingly another possibility, thus raising the intriguing possibility of multiple populations being present in M28. Coordinates, periods, and light curves in differential fluxes are provided for all the detected variables.}
|
\label{sec:intro} M28 (NGC~6266) is a moderately reddened [$E(B\!-\!V) = 0.40$~mag] and bright ($M_V = -8.16$~mag) globular cluster (GC) located at a low Galactic latitude ($b = -5.58$).\footnote{Unless otherwise noted, all cluster parameters in this paper are from \citet[][Dec. 2010 update]{wh96}.} Though relatively close, at a distance from the Sun of only 5.5~kpc, it remains a relatively ill-studied cluster, likely due to the unfavorable (and highly variable) foreground reddening. Indeed, variability studies of the cluster have so far been mostly restricted to photographic data \citep{wb90,wc90}, and modern color-magnitude diagrams (CMDs) have only recently been presented in the literature \citep*{tdea96,area00,vtea01,jagea12}. Still, M28 appears as a particularly interesting object for at least three reasons: first, it has been found to have a disk-like orbit \citep{rc91,ch93}, thus making it one of the most metal-poor members of the so-called ``thick disk'' family of GCs \citep[see, e.g.,][]{az88}. Second, among clusters of similar metallicity (${\rm [Fe/H]} \simeq -1.32$), M28 stands out as having a horizontal branch (HB) morphology strongly skewed towards the blue \citep[see also][]{ga81}, thus making it a ``second parameter'' cluster. The latter, if interpreted in terms of age, would accordingly indicate that the metal-poor tail of the thick disk is at least as old as the oldest components of the Galactic halo. Third, the cluster has been classified into an \citet{oo39,oo44} I (OoI) type, which is consistent with the relatively high metallicity of the cluster but in conflict with its HB type: with rare exceptions, blue HB clusters are commonly associated with type OoII \citep{rcea05,rcea10}. According to the \citet{ccea01} catalog,\footnote{\tiny\tt http://www.astro.utoronto.ca/$\sim$cclement/read.html} there are at present 24 variable stars known in the field of the cluster \citep[in addition to a millisecond pulsar;][]{alea87}. Of these, V1-V16 were discovered by \citet{hs49}, whereas V17-V24 were first discovered/reported on by \citet{wsh82,wsh84}. Updated ephemerides for these stars were provided in \citet{wb90} and \citet{wc90}, which remain the most recent papers to deal with the (photographic) light curves of M28 variable stars in a systematic way. \citet{rc91} reported, again based on photographic plates, on three additional candidate variables in the direction of the cluster, which however have not yet been systematically studied, or even incorporated into the electronic version of the \citeauthor{ccea01} catalog. There are strong reasons to believe that the current variable star tally for M28 is incomplete. First, while M28 is a very concentrated cluster, the photographic material which has been used in all previous variability studies does not allow the cluster core to be reliably resolved. Second, modern image-subtraction techniques \citep[e.g.,][]{al98,ca00,db08}, when applied to modern CCD images, have recently been providing rich harvests of variable stars towards the centers of even the previously best studied globular clusters \citep*[see, e.g.,][for recent examples]{rcea10,ckea11}. Accordingly, the main purpose of this paper is to provide the first CCD-based time-series photometry for the central regions of M28, in order to search for additional variables that may have gone unnoticed in previous studies and to provided updated periods and light curves for previously studied variables. We begin in \S2 by describing out dataset and reduction techniques. In \S3, we describe our variability results. A summary of our results is finally provided in \S4.
|
\label{sec:sum} We have presented the results of a new search for variable stars in M28. Our search has led to the discovery of a number of previously unknown variables, most of which are c-type RR Lyrae stars. The properties of the ab-type RR Lyrae stars are most consistent with an Oosterhoff-intermediate classification, but this is not clearly supported by the properties of the c-type RR Lyrae stars. A ``hybrid'' Oosterhoff I/II classification is thus possible, raising the question as to whether multiple populations may be present in this fairly massive cluster. More extensive, higher-quality datasets will be required to put the properties of this cluster on a firmer basis.
| 12
| 6
|
1206.4565
|
1206
|
1206.3173_arXiv.txt
|
Convection in the solar interior is thought to comprise structures on a spectrum of scales. This conclusion emerges from phenomenological studies and numerical simulations, though neither covers the proper range of dynamical parameters of solar convection. Here, we analyze observations of the wavefield in the solar photosphere using techniques of time-distance helioseismology to image flows in the solar interior. We downsample and synthesize 900 billion wavefield observations to produce 3 billion cross-correlations, which we average and fit, measuring 5 million wave travel times. Using these travel times, we deduce the underlying flow systems and study their statistics to bound convective velocity magnitudes in the solar interior, as a function of depth and spherical-harmonic degree $\ell$. Within the wavenumber band $\ell<60$, Convective velocities are 20-100 times weaker than current theoretical estimates. This suggests the prevalence of a different paradigm of turbulence from that predicted by existing models, prompting the question: what mechanism transports the heat flux of a solar luminosity outwards? Advection is dominated by Coriolis forces for wavenumbers $\ell<60$, with Rossby numbers smaller than $\sim10^{-2}$ at $r/R_\odot=0.96$, suggesting that the Sun may be a much faster rotator than previously thought, and that large-scale convection may be quasi-geostrophic. The fact that iso-rotation contours in the Sun are not co-aligned with the axis of rotation suggests the presence of a latitudinal entropy gradient.
|
The thin photosphere of the Sun, where thermal transport is dominated by free-streaming radiation, shows a spectrum in which granulation and supergranulation are most prominent. Observed properties of granules, such as spatial scales, radiative intensity and photospheric spectral-line formation are successfully reproduced by numerical simulations \cite{stein00,voegler2005}. In contrast, convection in the interior is not directly observable and likely governed by aspects more difficult to model, such as the integrity of descending plumes to diffusion and various instabilities \cite{rast98}. Further, solar convection is governed by extreme parameters \cite{miesch05} (Prandtl number $\sim 10^{-6} - 10^{-4}$, Rayleigh number $\sim 10^{19} - 10^{24}$, and Reynolds number $\sim 10^{12} - 10^{16}$), which make fully resolved three-dimensional direct numerical simulations impossible for the foreseeable future. It is likewise difficult to reproduce them in laboratory experiments. Turning to phenomenology, mixing-length theory (MLT) is predicated on the assumption that parcels of fluid of a specified spatial and velocity scale transport heat over one length scale (termed the {\it mixing length}) and are then mixed in the new environment. While this picture is simplistic \cite{cox_2004}, it has been successful in predicting aspects of solar structure as well as the dominant scale and magnitude of observed surface velocities. MLT posits a spatial convective scale that increases with depth (while velocities reduce) and coherent large scales of convection, termed {\it giant cells}. Simulations of anelastic global convection \cite{miesch_etal_08,charbonneau10,kapyla1, kapyla2}, more sophisticated than MLT, support the classical picture of a turbulent cascade. The ASH simulations \cite{miesch_etal_08} solve the non-linear compressible Navier-Stokes equations in the anelastic limit, i.e., where acoustic waves, which oscillate at very different timescales, are filtered out. Considerable effort has been spent in attempting surface \cite{hathaway00} and interior detection \cite{duvall, duvall03} of giant cells, but evidence supporting their existence has remained inconclusive.
|
\subsection{Convective transport} The spectral distribution of power due to an ensemble of convective structures, of spatial sizes small or large or both, will be broad. For example, it has been argued (\cite{hathaway00}) that photospheric convection comprises only granules and supergranules, and that the power spectrum of an ensemble of these structures would extend from the lowest to highest $\ell$. In other words, if granulation-related flow velocities were to be altered, the {\it entire} power spectrum would be affected. Thus the large scales which we image here (i.e., power for low $\ell$), contain contributions from small and large structures alike, and represent, albeit in a complicated and incomplete manner, gross features of the transport mechanism. Our constraints show that for wavenumbers $\ell < 60$, flow velocities associated with solar convection ($r/R_\odot = 0.96$) are substantially smaller than current predictions. Alternately one may interpret the constraints as a statement that the temporal coherence of convective structures is substantially shorter than predicted by current theories. Analysis of numerical simulations (\cite{miesch_etal_08}) of solar convection shows that a dominant fraction ($\sim 80$\%) of the heat transport is effected by the small scales, However, our observations show that the simulated velocities are substantially over-estimated in the wavenumber band $\ell < 60$, placing in question (based on the preceding argument) the entire predicted spectrum of convective flows and the conclusions derived thereof. % We further state that we lack definitive knowledge on the energy-carrying scales in the convection zone. We may thus ask: how would this paradigm of turbulence affect extant theories of dynamo action? For example, consider the scenario discussed by \cite{Spruit97}, who envisaged very weak upflows, which, seeded at the base of the convection zone, grow to ever larger scales due to the decreasing density as they buoyantly rise. These flows are in mass balance with cool inter-granular plumes which, formed at the photosphere, are squeezed ever more so as they plunge into the interior. Such a mechanism presupposes that these descending plumes fall nearly ballistically through the convection zone, almost as if a cold sleet, amid warm upwardly diffusing plasma. In this schema, individual structures associated with the transport process would elude detection because the upflows would be too weak and the downflows of too small a structural size (M. Sch\"{u}ssler, private communication, 2011). When viewed in terms of spherical harmonics, the associated velocities at large scales (i.e., low $\ell$), which contain contributions from both upflows and descending plumes, would also be small. Whatever mechanism may prevail, the stability of descending plumes at high Rayleigh and Reynolds numbers and very low Prandtl number is likely to play a central role (\cite{Spruit97, rast98}). \subsection{Differential Rotation and Meridional Circulation} Differential rotation, a large-scale feature ($\ell \sim 2$), is one individual global flow system and easily detected in our travel-time maps. Differential rotation is the only feature we ``detect" within this wavenumber band. In other words, upon subtracting this $\ell =2$ feature from the travel-time maps, the variance of the remnant falls roughly as $T^{-1}$, where $T$ is the temporal averaging length, suggesting the non-existence of other structures at these scales. Consequently, we may assert that we do not see evidence for a ``classical" inverse cascade that results in the production of a smooth distribution of scales. Current models of solar dynamo action posit that differential rotation drives the process of converting poloidal to toroidal flux. This would result in a continuous loss of energy from the differentially rotating convective envelope and Reynolds' stresses have long been thought of as a means to replenish and sustain the angular velocity gradient. The low Rossby numbers in our observations indicate that turbulence is geostrophically arranged over wavenumbers $\ell < 60$ at the depth $r/R_\odot = 0.96$, further implying very weak Reynolds stresses. Because flow velocities are likely to become weaker with depth in the convection zone, the Rossby numbers will decrease correspondingly. At wavenumbers of $\ell \sim 2$, the thermal wind balance equation describing geostrophic turbulence likely holds extremely well within most of the convection zone: \begin{equation} \Omega_0 \frac{\partial\Omega}{\partial z} = \frac{C}{r^2\sin\theta}\frac{\partial S}{\partial\theta}, \label{geostrophy} \end{equation} where $\Omega_0$ is the mean solar rotation rate, $\Omega$ is the differential rotation, $z$ is the axis of rotation, $\theta$ is the latitude, $C$ is a constant, $S$ is the azimuthally and temporally averaged entropy gradient. Differential rotation around $\ell \sim 2$ is helioseismically well constrained, i.e., the left side of equation~(\ref{geostrophy}) is accurately known (e.g., \cite{kosovichev97}). The iso-rotation contours are not co-aligned with the axis of rotation, yielding a non-zero left side of equation~(\ref{geostrophy}). Taylor-Proudman balance is broken and we may reasonably infer that the Sun does indeed possess a latitudinal entropy gradient, of a suitable form so as to sustain solar differential rotation (see e.g., \cite{kitchatinov95, balbus09}). The inferred weakness of Reynolds stresses poses a problem to theories of meridional circulation, which rely on the former to effect angular momentum transport in order to sustain the latter. Very weak turbulent stresses would imply a correspondingly weak meridional circulation (e.g., \cite{rempel2005}).%
| 12
| 6
|
1206.3173
|
1206
|
1206.1340_arXiv.txt
|
We investigate the spatial distribution of galactic satellites in high resolution simulations of structure formation in the $\Lambda$CDM model: the Aquarius dark matter simulations of individual halos and the Millennium II simulation of a large cosmological volume. To relate the simulations to observations of the Milky Way we use two alternative models to populate dark halos with ``visible'' galaxies: a semi-analytic model of galaxy formation and an abundance matching technique. We find that the radial density profile of massive satellites roughly follows that of the dark matter halo (unlike the distribution of dark matter subhalos). Furthermore, our two galaxy formation models give results consistent with the observed profile of the 11 classical satellites of the Milky Way. Our simulations predict that larger, fainter samples of satellites should still retain this profile at least up to samples of 100 satellites. The angular distribution of the classical satellites of the Milky Way is known to be highly anisotropic. Depending on the exact measure of flattening, 5--10 per cent of satellite systems in our simulations are as flat as the Milky Way's and this fraction does not change when we correct for possible obscuration of satellites by the Galactic disk. A moderate flattening of satellite systems is a general property of $\Lambda$CDM, best understood as the consequence of preferential accretion along filaments of the cosmic web. Accretion of a single rich group of satellites can enhance the flattening due to such anisotropic accretion. We verify that a typical Milky Way-mass CDM halo does not acquire its 11 most massive satellites from a small number of rich groups. Single--group accretion becomes more likely for less massive satellites. Our model predictions should be testable with forthcoming studies of satellite systems in other galaxies and surveys of fainter satellites in the Milky Way.
|
The satellites of the Milky Way offer a number of critical tests of the current cosmological paradigm, the $\Lambda$CDM model. Their abundance and internal structure are sensitive to the nature of the dark matter and thus to the high frequency end of the linear power spectrum of density fluctuations. Their spatial distribution is sensitive to the gravitational evolution of dark matter on galactic and supergalactic scales. High resolution simulations of the formation of galactic cold dark matter halos have revealed that a large number of substructures survive to the present day, accounting for about 10\% of the total halo mass \citep[][and references therein]{diemand07,springel08}. Since only about two dozen satellites are known to orbit in the halo of the Milky Way, this property is frequently deemed to pose a ``satellite problem'' for CDM-based cosmologies. In fact, it was shown about a decade ago that basic processes that are unavoidable during galaxy formation, such as supernova feedback and early Hydrogen reionization, readily explain why a visible satellite galaxy can form only in a tiny fraction of the surviving subhalos \citep{bullock00,benson2002,somerville2002}. This result has been confirmed repeatedly in recent years using semi-analytic models \citep{cooper10,guo11,li10,maccio10,font11}, cosmological gasdynamical simulations \citep{okamoto_frenk10,parry12,wadepuhl11} and simplified semi-empirical models \citep{kravtsov04,koposov09,busha10,munoz09}. A different kind of theoretical challenge is posed by the spatial distribution of the 11 classical satellites of the Milky Way. \cite{lyndenbell76}, \cite{kunkel79} and \cite{lyndenbell82} noted that these satellites lie very close to a great circle on the sky that is approximately perpendicular to the Galactic Plane. \cite{kroupa05} deemed such a highly flattened structure to be extremely unlikely in a CDM cosmology but, using N-body simulations of halo formation in the $\Lambda$CDM model, \cite{kang05}, \cite{libeskind05}, \cite{zentner05} and \cite{libeskind09} showed explicitly that this presumption is incorrect. Such flattened satellite distributions were dubbed ``great pancakes'' by \cite{libeskind05} who ascribed them to highly anisotropic accretion of proto-subhalos along the filaments of the cosmic web. Correlated accretion along filaments was also identified as the cause for the polar alignment of satellite disks found by \cite{deason11} in 20 percent of satellite systems (with more than 10 bright members each) in the {\small ``GIMIC''} N-body/gasdynamic simulations \citep{crain09}. Although all of these studies found that flattening of satellite systems is common in $\Lambda$CDM, they also found that the high degree of flattening in the Milky Way system is atypical. \cite{li08} showed that the flattening effects of anisotropic accretion are greatly enhanced in cases where infalling dark matter subhalos belong to a `group' sharing similar infall times and orbital angular momentum orientations (the members of such groups do not have to be bound in a single DM halo before infall). In a Mikly Way-mass system, they found that samples of $\sim10$ subhalos drawn from one or two such groups readily produced configurations as flat as that of the classical Milky Way satellites. Extrapolating this result, they suggested that the highly correlated infall of a rich group of satellites was a {\em possible} explanation for the abnormal degree of flattening seen in the Milky Way. However, they could not say whether this was a {\em probable} explanation since they did not calculate the likelihood of all the 11 {\em brightest} satellites being members of only one or two such groups. An important limitation of the simulations that have been analysed so far to study the spatial distribution of satellites is their relatively low resolution. Low resolution can give rise to excessive tidal disruption and to the artificial merging of some subhalos, potentially obscuring the true spatial distribution. In this paper, we use the suite of six simulations of individual Galactic halos from the Aquarius project, which are amongst the highest resolution CDM simulations carried out to date \citep{springel08}, as well as a sample of similar halos from the Millennium-II simulation \citep{boylan09}. Millennium-II has lower resolution than Aquarius but follows halo formation in a cosmological volume (a cube of side 100 $h^{-1}$Mpc, where $h$ denotes the Hubble constant in units of 100 ${\rm km s}^{-1}{\rm Mpc}^{-1}$). With the Aquarius simulations we are able to study satellites down to very small masses in six galactic halos and with the Millennium-II we are able to study the massive satellites of a large sample of galactic halos. Since the Aquarius halos are zoom-resimulations of regions in the Millennium-II volume, we are able also to study the effects of numerical resolution. We rank satellites in our simulations by stellar mass using two different techniques: semi-analytic modelling and a simple assignment of the brightest satellites to the most massive proto-subhalos. As we show, the two approaches pick out similar subsets of subhalos as satellite hosts. With these samples, we revisit the radial distribution of satellites and the great pancake and, for the first time, we investigate how these properties depend on the number of satellites considered (in samples ranked by stellar mass). This allows us to make predictions for forthcoming surveys such as Pan-STARRS \citep{kaiser10}, which may discover a large population of very faint satellites in the Milky Way and M31, and the ongoing Pandas survey of M31 \citep{mcconnachie09, martin09}. We also extend the work of \citet{li08} by using Aquarius to investigate the multiplicity function of groups of massive satellites. The paper is organized as follows. In Section~\ref{sec:model}, we discuss the suite of simulations and galaxy formation models that we use. In Section~\ref{sec:results}, we investigate the radial distribution of satellites, the prevalence of great pancakes, and groups of massive satellites. Our conclusions and discussion are presented in Section~\ref{sec:conclusion}. \begin{figure} \bc \hspace{1.cm} \resizebox{10cm}{!}{\includegraphics{figures/Fig1.ps}}\\% \caption{The relationship between the stellar mass assigned by the \citet{cooper10} semi-analytic model to subhalos in the six Aquarius simulations and the maximum value of the rotation curve attained by each subhalo over its entire past history. (Satellites with unresolved dark matter haloes are excluded.) The tight correlation for the most massive satellites motivates a simple $V_{\rm peak}$ model for ranking satellites by stellar mass.} \label{fig:Fig1} \ec \end{figure} \begin{figure} \bc \hspace{1.cm}% \resizebox{10cm}{!}{\includegraphics{figures/Fig2.ps}}\\% \caption{The projected positions of satellites identified using the \citet{cooper10} semi-analytic model (diamonds) and the $V_{\rm peak}$ method (filled circles) in the Aq-A2 halo. From top left to bottom right, positions are shown for the top 11, 30, 60 and 100 satellites ranked by stellar mass in each model. The overlap ratio between the two samples is given in the legend. It decreases from $f=0.91$ for the top 11 to $f=0.7$ for the top 100 satellites.} \label{Fig2} \ec \end{figure}
|
\label{sec:conclusion} The spatial distribution of satellites in the Milky Way and other galaxies reflects both the nature of the dark matter and the processes of galaxy formation. In this paper we have used high mass resolution cosmological simulations of structure formation in the $\Lambda$CDM model: the Aquarius resimulations of galactic scale dark matter halos and the Millennium II simulation (MRII) of a cubic volume 100 $h^{-1}$Mpc on a side, with about 2000 halos of mass comparable to the Milky Way's halo. We have further employed two models of galaxy formation to trace satellites in the simulation: the semi-analytic models of \cite{cooper10} and \cite{guo11} for Aquarius and MRII respectively and a less sophisticated model in which galaxy stellar mass is assumed to be proportional to $V_{\rm peak}$, the highest circular velocity attained by the halo throughout its life. Our two methods place 10 of the most massive 11 satellites in the same Aquarius halos. The combination of high-resolution simulations and robust galaxy formation models allows us to identify satellites in a reliable way. We find that good mass resolution is essential to obtain an accurate estimate of the radial distribution of satellites. Indeed, even at the resolution of the MRII, $m_p=6.9 \times 10^6 h^{-1} M_{\odot}$, many genuine substructures are artificially destroyed, as may be seen in Fig.~\ref{fig:Fig5}. In this case, the $V_{\rm peak}$ method for assigning stellar mass to satellites, as well as all previous cosmological simulations which have poorer resolution \citep[e.g.][]{libeskind05,kang05}, would give inaccurate results. By contrast, our semi-analytic model, which continues to track satellites even after they have lost their halos does give a faithful prediction of the theoretical expectations in the $\Lambda$CDM model. In the case of the Aquarius simulations, the resolution is good enough that galaxies with unresolved halos galaxies are unimportant \citep{font11} and we can compare our semi-analytic results with the $V_{\rm peak}$ method. The results from the two ranking methods agree well. Our main results concerning the satellite radial distribution are: (i) there is substantial halo-to-halo scatter in the radial distribution of satellites in the Aquarius simulations, but in all cases the massive satellites have a similar radial density profile to that of the dark matter; (ii) both the semi-analytic and $V_{\rm peak}$ models give results in good agreement with the measured radial distribution of the 11 classical satellites of the Milky Way; (iii) the radial density profile of larger satellite samples, going down to lower masses, is similar to that of the 11 most massive, at least up to samples of 100 satellites. This prediction may be tested by future surveys such as Pan-STARRS. We also investigated the angular distribution of satellites. This is interesting because the Milky Way's 11 brightest satellites show a very anisotropic distribution, the ``Great pancake'' of \cite{libeskind05}. We characterized the angular distribution by measuring both the axial ratios of the satellite system and the thickness of the slab in which the satellites are concentrated. We verified that the peculiar distribution seen in the Milky Way is highly significant: only 1 percent of isotropic samples with the same radial distribution are flatter than the Milky Way's system. In the 1686 halos in the MRII which have a mass plausibly similar to that of the Milky Way we find systems flatter than that of the classical satellites with probabilities of 6 and 13 percent according to the axial ratio and slab thickness tests respectively. These probabilities are slightly lower than that found by \citet{libeskind05}, who used a slightly incorrect value for the flattening of the Milky Way satellites. Our probabilities are also lower than those found by by \citet{kang05}, a discrepancy that is readily understood as a consequence of the poor mass resolution of the simulations used by these authors. The Milky Way system thus appears to be in the tail of the flattening distribution expected for massive satellites in the $\Lambda$CDM model. The presence of a satellite as massive as the LMC \citep{boylan10,busha11,guoq11,liu11} and the polar alignment of the system \citep{deason11} seem to be comparably rare, but still consistent with the predicted distributions. Finally, we investigated the extent to which satellites are accreted in groups. This is interesting for several reasons, including the possibility that the Great pancake might be explained by multiple accretion in 2-3 groups \citep{li08}. Our simulations confirm that this is a rare occurrence in $\Lambda$CDM. On average, only 30 percent of the top 11 satellites in the Aquarius simulations share a friends-of-friends group with another top 11 satellite before infall; the rest come into the main halo without any companions of comparable mass. However, multiple accretion becomes increasingly important for larger, fainter samples of satellites. For example, in samples of the 60 most massive satellites in two of the Aquarius halos (Aq-B and Aq-F), we find that as many as 20 come into the main halo in a single group and as many as 11 in the other simulated halos. This interesting property may potentially have observational consequences \citep[e.g.][]{sales12}. \cite{libeskind05} proposed that filamentary accretion is responsible for the flattening of Milky Way-like satellite populations. \cite{lovell11} subsequently showed that the subhalos of the Aquarius simulations have strongly correlated orbital angular momenta as the result of anisotropic accretion. Recently, however, \citet{pawlowski12} re-analysed the angular momentum orientations of subhalos in Aquarius (as presented in \citealt{lovell11}) and concluded that anisotropic accretion is unimportant for producing flattened satellite systems, in direct contradiction with the results of \cite{libeskind05} and \cite{lovell11}. They required that the Milky Way flattening should be reproduced in the mean, rather than being merely consistent with the expected distribution. As the Milky Way is in the $\sim10$ per cent tail of the distribution predicted by $\Lambda$CDM (as our study and others have shown), they reject a CDM model in favour of tidal galaxy formation, which they claim readily produces highly correlated disks of satellites. We believe that the explanation offered by \cite{libeskind05} remains the most appropriate. A single randomly chosen system such as the Milky Way is extremely unlikely perfectly to represent the mean value of every measurable property. A larger sample of satellites around other galaxies will test the tidal formation hypothesis of \citet{pawlowski12} in which highly flattened configurations are easily achieved and should therefore be the norm. If, on the other hand, the CDM model is a realistic description of nature, then the average satellite configurations should be only moderately flattened, as illustrated in Figs.~\ref{fig:Fig7} and~\ref{fig:Fig8}. Our simulations are useful not only to test our models against data for the classical satellites of the Milky Way, as we have done here, but also to make predictions for future surveys that will quantify the spatial distribution of larger and fainter satellite samples, both in the Milky Way and in other galaxies. For a sample of Milky Way satellites complete to very faint magnitudes, we have shown that the luminosity function \citep{font11}, radial distribution and system shape should vary less from halo to halo than is the case for the most massive eleven. Measuring these three highly characteristic properties of the satellite system test both the $\Lambda$CDM cosmology and models of galaxy formation in novel, interesting ways.
| 12
| 6
|
1206.1340
|
1206
|
1206.4938_arXiv.txt
|
We report the discovery of a unique object, BD+48 740, a lithium overabundant giant with A(Li)=2.33 $\pm$ 0.04 (where A(Li)=$\log n_{Li}/n_{H}+12$), that exhibits radial velocity (RV) variations consistent with a 1.6 M$_{J}$ companion in a highly eccentric, e=0.67 $\pm$ 0.17 and extended, a=1.89~AU (P=771~d), orbit. The high eccentricity of the planet is uncommon among planetary systems orbiting evolved stars and so is the high lithium abundance in a giant star. The ingestion by the star of a putative second planet in the system originally in a closer orbit, could possibly allow for a single explanation to these two exceptional facts. If the planet candidate is confirmed by future RV observations, it might represent the first example of the remnant of a multiple planetary system possibly affected by stellar evolution.
|
There has been a growing number of exoplanets detected around post-main sequence (MS) stars. At present, about 50 red giants (RGs) are known to host planetary or brown dwarf-mass companions. Recent discoveries also show evidence of multi--planet systems (cf. HD 102272 b, c) or multi - brown dwarf systems (BD+20 2457 b,c -- \citealt{Niedzielski2009b}). Although hot jupiter type planets exist around sub--giants \citep{Johnson2010}, the more evolved giants show a paucity of short period and eccentric planets \citep{Johnson2007}. This is most likely the result of tidal disruption and/or planet engulfment during the RG phase (e.g.\citealt{VillaLivio2009}). Even rarer is the number of lithium-rich RG stars. According to the standard evolution theory, when a solar--type star leaves the MS and becomes a RG, its lithium abundance should drop from A(Li) $\sim 3.3$ to about a 1.5--level. In fact, the observed upper giant branch limit is A(Li) $<$ 0.5 \citep{LindPrimas2009} and only a few percent of giants have been observed to have A(Li) $>$ 1.5 (\cite{Kumar2011} and references therein), with some exhibiting A(Li) $\sim$ 3.3, the value expected for a protostellar disk rather than an evolved star. In this paper, we report the discovery of a planet in a highly eccentric, long-period orbit around a lithium-overabundant giant star.
|
We present evidence of a high lithium abundance, A(Li)=2.33 $\pm$ 0.04, in the RG star BD+48 740. This value is well above the expected limit of A(Li)=1.5 for evolved stars. We also present multi-epoch, precise radial velocities for the star which, although sparse, show periodic variations which can be interpreted as a result of the Keplerian motion of a planetary mass companion with $m\sin i=1.6$ M$_{J}$, on a highly eccentric orbit of e=0.67. We find no evidence of a stellar-mass companion to BD+48 740. Given the current evolutionary status of the star, its Li abundance and the planet's current orbit, we discuss a possibility that BD+48 740 had a second planet in an innermost orbit that could have been engulfed by the star. This possibility, although not directly verifiable with the available data, allows a scenario in which both the high lithium abundance of the star and the planet's highly eccentric orbit could originate from a recent, violent dynamical event, likely representing a later stage in the evolution of the planetary system, which was similar to HAT P-13 while still on the MS. BD+48 740 is the first example of Li-overabundant giant star with a planetary mass companion candidate.
| 12
| 6
|
1206.4938
|
1206
|
1206.3729_arXiv.txt
|
{The ACME Spectra project provides absolutely calibrated, mostly empirical spectra of exoplanet host stars for use in analysis of the stars and their planets. Spectra are obtained from ground-based telescopes and are tied directly to calibrated ground- and space-based photometry. The spectra remain only ``mostly'' empirical because of telluric absorption, but interpolation of stellar models over the gaps in wavelength coverage provides continuous stellar spectra. Among other uses, the spectra are suitable for precisely converting observed secondary eclipses (occultations) into absolute flux units with minimal recourse to models. In this letter I introduce ACME's methods and present a calibrated spectrum of the nearby, super-Earth hosting star 55~Cancri that spans the range from 0.81--5.05\,\micron. This spectrum is well-suited for interpreting near- and thermal-infrared eclipse observations. With this spectrum I show that the brightness temperature of the small, low-mass transiting planet 55~Cnc~e is \tbright\ at 4.5\,\micron\ (cooler than previously reported), which corresponds to a planetary flux of \pflux. This result suggests the planet has some combination of a nonzero albedo, a moderately efficient redistribution of absorbed stellar irradiation, and/or an optically thick atmosphere, but more precise eclipse measurements are required to distinguish between these scenarii.} { } { } { } { }
|
Measurements of flux densities are the primary diagnostic of conditions in stellar and planetary atmospheres, and are used to constrain albedos, sizes, temperatures abundances, temperature profiles, and even interior conditions of stars and planets. In the last decade observations of transiting planets during eclipse have begun to reveal the intrinsic emission of these externally irradiated bodies \citep[e.g.,][]{deming:2005,charbonneau:2005}. Though initially confined to space-based observatories such as Spitzer \citep[and, more recently, Hubble;][]{swain:2012}, more favorable targets and improved observing strategies have recently enabled a growing number of ground-based measurements \citep[e.g.,][]{sing:2009ogle,demooij:2009}. Because eclipse measurements are inherently relative, precise planetary fluxes require stellar fluxes yet more precise. The state of the art is now reaching the point at which this condition is no longer met; for example, recent measurements of eclipses with Spitzer/IRAC have attained relative precisions of $1-4\%$ \citep[e.g.,][]{agol:2010, campo:2011}, comparable to the instrument's 3\% absolute calibration accuracy \citep{SSC:2012}. The goal of the ACME project is to provide a catalog of Absolutely Calibrated, Mostly Empirical (ACME) spectra of exoplanet host stars with high accuracy and precision. In the rapidly approaching era of exoplanet spectroscopy \citep[e.g.,][]{swain:2012} it is imperative that such an effort advance beyond standard photometric calibrations and focus on calibrated spectroscopy. In addition to the aforementioned conversion of planetary eclipse depths into absolute flux units, other potential uses for such a catalog include the study of the planets' host stars: abundance analyses, spectral typing, and refined bolometric luminosities and effective temperatures are all enabled by such data. I describe the observation, reduction, and calibration of ACME spectra in \S~\ref{sec:obs}. As a particular case study I present a calibrated spectrum of the nearby star \object{55~Cancri} (\object{55~Cnc}) in \S~\ref{sec:results}\footnote{Spectral data are available in electronic form at the CDS via anonymous ftp to cdsarc.u-strasbg.fr (130.79.128.5) or via \url{http://cdsweb.u-strasbg.fr/cgi-bin/qcat?J/A+A/} }. With this calibrated spectrum I provide an improved measurement of \object{55~Cnc~e}'s flux at 4.5\,\micron\ in light of recent observations \citep{demory:2012}, and I conclude with future prospects.
| 12
| 6
|
1206.3729
|
|
1206
|
1206.5339_arXiv.txt
|
{The classic HII region M17 is one of the best studied across the electromagnetic spectrum. We present sensitive, high angular resolution observations made with the Jansky Very Large Array (JVLA) at 4.96 and 8.46 GHz that reveal the presence of 38 compact radio sources, in addition to the well known hypercompact cometary HII region M17 UC1. For this last source we find that its spectral index of value $\sim$1 is due to a gradient in opacity across its face. Of the 38 compact radio sources detected, 19 have stellar counterparts detected in the infrared, optical, or X-rays. Finally, we discuss the nature of the radio emission from the massive binary system CEN 1a and 1b, concluding that both are most probably non-thermal emitters, although the first is strongly time variable and the second is steady.}
|
The centimeter continuum radiation from classic HII regions is dominated by strong free-free emission from the extended ionized gas present there. However, when observed with the high angular resolution provided by an interferometer, the extended emission is filtered out and one starts to detect compact sub-arcsecond sources of various natures (see Garay et al. 1987; Churchwell et al. 1987; Felli et al. 1993; Zapata et al. 2004 for the case of Orion A). The brightest of these sources are the hypercompact (HC) HII regions, that trace the ionized gas produced by young OB stars still embedded in dense molecular gas (e.g. Kurtz 2005; Lizano 2008). The externally ionized globules are also sources of free-free radiation and result from the interaction of the UV photons of the OB stars in the region with remaining blobs of neutral gas existing inside the HII region (e.g. Garay et al. 1987). The proplyds (O'Dell et al. 1993) are similar to the externally ionized globules, but in this case the object being ionized is a protoplanetary disk surrounding a young star. The last two known types of free-free emitters are the jets emanating from accreting protostars (Anglada 1996; Eisloffel et al. 2000) and the spherical ionized winds produced by massive stars (e.g. Bieging et al. 1989; Contreras et al. 1996). There are also two types of sources with non-thermal radio continuum emission. Young low-mass stars can have strong magnetospheric activity and emit detectable gyrosynchrotron radiation (Andr\'e et al. 1988). Finally, there is also strong evidence that systems of massive binary stars can produce synchrotron radiation in the region where their winds collide (e.g. Pittard et al. 2006; Ortiz-Le\'on et al. 2011). In Table 1 we present a summary of the characteristics of these different types of compact radio sources. In this paper we present a sensitive, high angular resolution study made with the Jansky Very Large Array (JVLA) of the National Radio Astronomy Observatory (NRAO)\footnote{The NRAO is operated by Associated Universities Inc. under cooperative agreement with the National Science Foundation.} toward the classic HII region M17 (the Omega Nebula, NGC~6618, W38, S45). In \S~2 we present the observations, while in \S~3 we list and briefly discuss the compact radio sources detected. In \S~4 we use our data to present an explanation for the spectral index of order 1 observed in the hypercompact HII region M17 UC1. In \S~5 we discuss the time variable sources in our sample, while in \S~6 we concentrate on CEN 1a and 1b, the members of the massive binary system that ionizes most of M17. In \S~7 we try to model the time-variable emission of CEN 1a in terms of a thermal model, concluding that this is not feasible. Finally, \S~8 presents a brief discussion on some of the other individual sources and in \S~9 we summarize our conclusions.
|
We presented sensitive, high angular resolution observations made with the Jansky Very Large Array (JVLA) at 4.96 and 8.46 GHz toward the HII region M17. Our main conclusions are listed in what follows. 1. We detected 38 compact radio sources, practically all associated with the region. Only 19 of these sources have counterparts, in all cases stars previously detected at infrared, optical, or X-rays. We argue that the radio sources without counterparts are similar to those with counterparts, but that they are heavily obscured, which makes detection difficult at wavelengths other than radio. 2. We studied the spectral index of the hypercompact HII region M17 UC1, finding that its spectral index of value $\sim$1 is due to a gradient in optical depth across its face. We speculate that this type of gradient could explain similar spectral indices detected in other hypercompact HII regions. In this case, all these sources should show a cometary morphology. 3. We discussed the nature of the radio emission from the massive binary system CEN 1a and 1b, concluding that both are most probably non-thermal emitters, although the first one is highly time variable and the second one is steady. Non-thermal emission in massive stars is usually attributed to them being binary systems, with the emission coming from the wind collision region. This system provides the opportunity to study simultaneously two stars, a variable one (CEN 1a) and a steady one (CEN 1b) and to try to understand the reasons for their different behaviours.
| 12
| 6
|
1206.5339
|
1206
|
1206.5755_arXiv.txt
|
Sylvia is a triple asteroid system located in the main belt. We report new adaptive optics observations of this system that extend the baseline of existing astrometric observations to a decade. We present the first fully dynamical 3-body model for this system by fitting to all available astrometric measurements. This model simultaneously fits for individual masses, orbits, and primary oblateness. We find that Sylvia is composed of a dominant central mass surrounded by two satellites orbiting at 706.5 $\pm$ 2.5 km and 1357 $\pm$ 4.0 km, i.e.,\ about 5 and nearly 10 primary radii. We derive individual masses of 1.484$_{-0.014}^{+0.016}$ $\times 10^{19}$ kg for the primary (corresponding to a density of 1.29 $\pm$ 0.39 g cm$^{-3}$), 7.33$_{-2.3}^{+4.7}$ $\times 10^{14}$ kg for the inner satellite, and 9.32$_{-8.3}^{+20.7}$ $\times 10^{14}$ kg for the outer satellite. The oblateness of the primary induces substantial precession and the $J_2$ value can be constrained to the range of 0.0985$-$0.1. The orbits of the satellites are relatively circular with eccentricities less than 0.04. The spin axis of the primary body and the orbital poles of both satellites are all aligned within about two degrees of each other, indicating a nearly coplanar configuration and suggestive of satellite formation in or near the equatorial plane of the primary. We also investigate the past orbital evolution of the system by simulating the effects of a recent passage through 3:1 mean-motion eccentricity-type resonances. In some scenarios this allow us to place constraints on interior structure and past eccentricities.
|
\label{introduction} (87) Sylvia is a triple asteroid residing in the main belt, with heliocentric semi-major axis 3.5 AU, eccentricity 0.085, and inclination 11$^{\circ}$ relative to the ecliptic. Sylvia's outer satellite, named Romulus, was discovered in 2001 using the W. M. Keck Telescope \citep{brow01,marg01}, and was also detected in Hubble Space Telescope (HST) images \citep{stor01}. The inner satellite, Remus, was not discovered until the advent of improved adaptive optics systems in 2004 using the European Southern Observatory's Very Large Telescope (VLT) \citep{marc05}. The diameter of the primary has been estimated at $\sim$280 km through shape fits to adaptive optics images \citep{marc05}; this estimate is consistent with stellar occultation observations \citep{lin09}. Assuming this primary size, approximate sizes for the individual satellites have been estimated by adopting the same albedo as the primary and measuring each satellite's brightness relative to the primary. The diameter estimates are $\sim$7 km for Remus and $\sim$18 km for Romulus \citep{marc05}. Sylvia was the first triple asteroid system announced, even though the triple nature of 2002 CE26 was being actively discussed during the acquisition of the Sylvia observations~\citep{shep06}. Additional discoveries of multiples in the Solar System have followed. They include near-Earth triples (153591) 2001~SN263 \citep{nola08} and (136617) 1994~CC \citep{broz09}, main belt triples Kleopatra \citep{desc11}, Eugenia \citep{merl99,marc07}, Balam \citep{merl02,marc08}, and Minerva \citep{marc11}, and trans-Neptunian systems (47171) 1999~TC36 \citep{marg05,bene10}, Haumea \citep{brow05,brow06}, and the Pluto/Charon system \citep{weav06}. Following these discoveries, characterization of multiple systems have unearthed a wealth of information about their fundamental physical properties such as masses and densities, dynamical processes, and constraints on formation and evolutionary mechanisms. Such research has been possible because we can derive the masses of the individual components of a triple or higher-multiplicity system by analyzing their mutual gravitational interactions, which is possible in binary systems only when reflex motion is detected~\citep{marg02s, ostr06, naid11dps}. These masses in conjunction with size estimates can provide densities. Using this method, \citet{fang11} performed a detailed analysis of 2001~SN263 and 1994~CC, including masses, densities, and dynamical evolution. Similarly, work on the Pluto/Charon system and dwarf planet Haumea and its satellites have yielded information about their physical properties, tidal interactions, and evolutionary processes \citep{lee06,thol08,rago09}. The high scientific return from studies of binaries and triples has been reviewed by \citet{merl02} and \citet{noll08}. To date, no such dynamical orbit solution nor detailed analysis has been performed for Sylvia. Previous work by \citet{marc05} approximated the actual orbits of Remus and Romulus with individual two-body fits that included primary oblateness. However, drawbacks of such methods include the failure to account for third-body perturbations as well as the inability to solve for individual component masses. Additional researchers based their studies on the published two-body orbits \citep{marc05} plus unspecified component mass assumptions to study Sylvia's long-term evolution \citep{wint09,frou11}, even though component masses are undetermined and can span several orders of magnitude. In this work, we report additional Keck and VLT imaging data for Sylvia (Section \ref{observations}). Using primary$-$satellite separations measured from these data plus published astrometry, we present a fully dynamical 3-body orbital and mass solution for Sylvia, by accounting for mutually interacting orbits as well as the primary's non-sphericity (Section \ref{orbitsolution}). Although the orbital periods of the satellites are near a 8:3 ratio, we do not find that the system is currently in such a resonance (Section \ref{currentmmr}). We also analyze Sylvia's short-term and long-term future evolution (Section \ref{evolution}). Lastly, we investigate the past orbital evolution of Remus and Romulus by modeling passage through the 3:1 mean-motion resonance (Section \ref{origin}). A summary of main conclusions is given in Section \ref{conclusion}.
|
\label{conclusion} The goals of this study were to characterize Sylvia's current orbital configuration and masses as well as to illuminate the past orbital evolution of this system. Our work can be summarized as follows: (1) We reported new astrometric observations of Sylvia in 2011 that increased the number of existing epochs of astrometry by over 50\%. These new observations extended the existing baseline of observations to 7 years (for Remus) and to 10 years (for Romulus). (2) We fit a fully dynamical 3-body model to the available astrometric data. This model simultaneously solved for orbits of both satellites, individual masses, and the primary's oblateness (Table \ref{bestfit}). We found that the primary has a density of 1.29$\pm$0.39 g cm$^{-3}$ and is oblate with a $J_2$ value in the range of 0.0985$-$0.1. Constraints on satellite radii can be obtained from the mass determinations by assuming that the satellites have a bulk density equal to that of the primary; we find $\sim$4.5$-$6.1 km for Remus and $\sim$2.6$-$8.2 km for Romulus. These ranges would have to be modified if the actual density of the primary or of the satellites was different from the nominal value assumed here. The orbits of the satellites are relatively circular. We find that the primary's spin pole is best fit when aligned to Romulus' orbital pole, and that the satellites' orbit poles are coplanar to within one degree. (3) We numerically investigated the short-term and long-term stability of the orbits of Sylvia's satellites. There are periodic fluctuations in eccentricity for both satellites, most notably for the inner satellite Remus. We verified that these eccentricity excursions are due to the effects of primary oblateness. From long-term integrations we found that the system is in a very stable configuration, in agreement with previous investigations. (4) We studied the past orbital evolution of Sylvia's satellites, including the most recent low-order MMR resonance crossing, which is the 3:1. We used direct N-body integrations with forced tidal migration to model such an encounter. To examine the case of resonant capture followed by escape, we calculate the tidal damping timescale to go from the post-encounter eccentricity to the observed value. Using available tidal models, we find that the damping timescale for Romulus can be prohibitively large if its post-resonance eccentricity exceeded $\sim$0.023. This suggests that the system crossed the $e_2^2$ and $e_1e_2$ resonances without capture, or that it was not captured in these resonances for a sufficient duration to substantially increase the eccentricity of Romulus. Similar timescale constraints from tidal damping also imply that Remus may have a rubble pile structure if its post-resonance eccentricity exceeded $\sim$0.032. Alternatively, if no capture in any resonance occurred then we are able set lower limits on their past eccentricities ($e_1=0.00864$ and $e_2=0.00410$). The detailed characterization of Sylvia presented in this paper has allowed for analyses of its orbital evolution. Such studies of triple systems are important in order to understand their key physical properties, orbital architectures, and intriguing evolutionary histories.
| 12
| 6
|
1206.5755
|
1206
|
1206.0941_arXiv.txt
|
We present new constraints on the star formation histories of the ultra-faint dwarf (UFD) galaxies, using deep photometry obtained with the {\it Hubble Space Telescope (HST)}. A galaxy class recently discovered in the Sloan Digital Sky Survey, the UFDs appear to be an extension of the classical dwarf spheroidals to low luminosities, offering a new front in efforts to understand the missing satellite problem. They are the least luminous, most dark-matter dominated, and least chemically-evolved galaxies known. Our {\it HST} survey of six UFDs seeks to determine if these galaxies are true fossils from the early universe. We present here the preliminary analysis of three UFD galaxies: Hercules, Leo~IV, and Ursa Major~I. Classical dwarf spheroidals of the Local Group exhibit extended star formation histories, but these three Milky Way satellites are at least as old as the ancient globular cluster M92, with no evidence for intermediate-age populations. Their ages also appear to be synchronized to within $\sim$1~Gyr of each other, as might be expected if their star formation was truncated by a global event, such as reionization.
|
Although the Lambda Cold Dark Matter paradigm is consistent with many observable phenomena, discrepancies arise at small scales. One of the most prominent issues is that it predicts many more dark-matter halos than are actually seen as dwarf galaxies (e.g., Moore et al.\ 1999). A possible solution has arisen with the recent discovery of additional satellites around the Milky Way (e.g., Willman et al.\ 2005; Zucker et al.\ 2006; Belokurov et al.\ 2007) and Andromeda (e.g., Zucker et al.\ 2007) in the Sloan Digital Sky Survey (York et al.\ 2000) and other wide-field surveys (e.g., McConnachie et al.\ 2009). The newly-discovered ultra-faint dwarf (UFD) galaxies appear to be an extension of the classical dwarf spheroidals (dSphs) to lower luminosities ($M_V \gtrsim -8$~mag). UFD luminosities are comparable to those of globular clusters, but one distinction in the former is the presence of dark matter. Even massive globular clusters have mass-to-light ratios ($M/L_V$) of $\sim$2 (e.g., Baumgardt et al.\ 2009; van de Ven et al.\ 2006), precluding significant dark matter. In contrast, all known dwarf galaxies have higher $M/L_V$ (Kalirai et al.\ 2010 and references therein). UFD kinematics are clearly dark matter-dominated, with $M/L_V > 100$ (e.g., Kleyna et al.\ 2005; Simon \& Geha 2007; Mu\~noz et al.\ 2006), even where velocity dispersions have been revised downward (e.g., Koposov et al.\ 2011). The inferred dark-matter densities of dwarf galaxies suggest a high-redshift collapse for both classical dSphs and UFDs ($z \sim 12$; Strigari et al.\ 2008), but the dSphs apparently continued to evolve (Orban et al.\ 2008; Weisz et al.\ 2011). In contrast, the UFDs are the least chemically-evolved galaxies known, with abundance patterns that imply their star formation was brief (Frebel et al.\ 2010) and individual stellar metallicities as low as [Fe/H]~=~$-3.7$ (Norris et al.\ 2010). The strict conformance to a metallicity-luminosity relation for all Milky Way satellites limits the amount of tidal stripping to a factor of $\sim$3 in stellar mass (Kirby et al. 2011). Therefore, UFDs are not tidally stripped versions of classical dSphs (see also Penarrubia et al. 2008; Norris et al.\ 2010). \begin{table*}[t] \begin{center} \caption{Observations} \begin{tabular}{ccccccccccccc} \tableline & & & & & & & & & \multicolumn{2}{c}{Exposure per tile} & \multicolumn{2}{c}{50\% complete}\\ & R.A. & Dec. & $(m-M)_V$&$E(B-V)$&$M_V$ & &[Fe/H] & & F606W & F814W & F606W & F814W \\ Name & (J2000) & (J2000) & (mag) &(mag) &(mag) &$<$[Fe/H]$>$\tablenotemark{a}&r.m.s.\tablenotemark{b}&tiles& (s) & (s) & (mag) & (mag) \\ \tableline Hercules & 16:31:05&+12:47:07& 20.92$\pm$0.06 &0.08$\pm$0.02 &$-6.2$\tablenotemark{c}&$-2.41$ &0.6 & 2 & 12,880 & 12,745 & 29.1 & 29.1\\ Leo IV & 11:32:57&-00:31:00& 21.15$\pm$0.08 &0.05$\pm$0.02 &$-5.8$\tablenotemark{d}&$-2.54$ &0.9 & 1 & 20,530 & 20,530 & 29.1 & 29.2\\ Ursa Major I & 10:35:04&+51:56:51& 20.11$\pm$0.04 &0.04$\pm$0.02 &$-5.5$\tablenotemark{e}&$-2.18$ &0.7 & 9 & 4,215 & 3,725 & 28.4 & 28.4\\ \tableline \end{tabular} \end{center} \tablenotetext{a}{Kirby et al.\ (2011), based on Simon \& Geha et al.\ (2007) spectroscopy.} \tablenotetext{b}{For stars with $<$0.3~dex uncertainties.} \tablenotetext{c}{Sand et al.\ (2009)} \tablenotetext{d}{de Jong et al.\ (2010)} \tablenotetext{e}{Martin et al.\ (2008)} \end{table*} As one way of solving the missing satellite problem, galaxy formation simulations assume that UFDs formed the bulk of their stars prior to the epoch of reionization (e.g., Tumlinson 2010; Mu\~noz et al.\ 2009; Bovill \& Ricotti 2009; Koposov et al.\ 2009). Mechanisms that could drive an early termination of star formation include reionization, gas depletion, and supernova feedback. Using the {\it Hubble Space Telescope (HST)}, we are undertaking a deep imaging survey of UFDs that reaches the old main sequence (MS) in each galaxy, yielding high-precision color-magnitude diagrams (CMDs) that provide sensitive probes of their star formation histories. The program includes Hercules, Leo~IV, Ursa Major~I, Bootes~I, Coma Berenices, and Canes Venatici~II. Here, we give preliminary results for the first three galaxies.
|
The three UFD galaxies here are among the more distant of those known in the Milky Way system, at 100--150~kpc, so ground-based CMDs of each implied they were old ($>$10~Gyr), but could not put tight constraints on their ages (e.g., Sand et al.\ 2009; Sand et al.\ 2010; Okamoto et al.\ 2008, 2012; Ad\'en et al.\ 2010). Our {\it HST} observations reach well below the MS turnoff in each galaxy, revealing that all three host truly ancient metal-poor populations. The majority of the stars in each galaxy must have ages within 1~Gyr of M92's age, with younger ages strongly ruled out, although we cannot exclude a trace population of stars 1--2~Gyr younger than M92 at the high end of the metallicity distribution. Two UFD galaxies (Bootes~I and Coma Berenices) are significantly closer (44--66 kpc), such that ground-based CMDs can place constraints on their ages approaching those possible with {\it HST} (Mu\~noz et al.\ 2010; Okamoto et al.\ 2012). These CMDs also imply ages approximately as old as M92, although the use of distinct bandpasses hampers accurate comparisons to our observations. Our {\it HST} survey includes these galaxies, enabling accurate age measurements in a significant sample of UFDs, with all observed in the same photometric system as each other and the most ancient globular clusters. If we include Bootes~I and Coma Berenices, it seems likely that at least 5 UFD galaxies have ages consistent with that of the oldest known globular cluster (M92), with no evidence for significantly younger populations. This is in striking contrast to any other galaxy class in the local universe. Our external vantage point for Andromeda enables accurate ages throughout its halo, using the same instrument and techniques described here, yet all such measurements to date have found an extended star formation history, with significant numbers of stars younger than 10~Gyr (Brown et al.\ 2006, 2008). An {\it HST} survey of 60 dwarf galaxies within 4~Mpc found that most formed the bulk of their stars prior to $z \sim 1$, but none were consistent with a purely ancient population (Weisz et al.\ 2011). The UFDs may be the only galaxies where star formation ended in the earliest epoch of the universe. If so, the apparent synchronicity to their star formation histories suggests a truncation induced by a global event, such as reionization (13.3~Gyr ago; Jarosik et al.\ 2011). These UFDs were likely the victims of reionization, rather than the agents, given the small numbers of stars available to produce ionizing photons.
| 12
| 6
|
1206.0941
|
1206
|
1206.5878_arXiv.txt
|
We present a new parallel code for computing the dynamical evolution of collisional $N$-body systems with up to $N\sim10^{7}$ particles. Our code is based on the H\'enon Monte Carlo method for solving the Fokker-Planck equation, and makes assumptions of spherical symmetry and dynamical equilibrium. The principal algorithmic developments involve optimizing data structures, and the introduction of a parallel random number generation scheme, as well as a parallel sorting algorithm, required to find nearest neighbors for interactions and to compute the gravitational potential. The new algorithms we introduce along with our choice of decomposition scheme minimize communication costs and ensure optimal distribution of data and workload among the processing units. Our implementation uses the Message Passing Interface (MPI) library for communication, which makes it portable to many different supercomputing architectures. We validate the code by calculating the evolution of clusters with initial Plummer distribution functions up to core collapse with the number of stars, $N$, spanning three orders of magnitude, from $10^{5}$ to $10^{7}$. We find that our results are in good agreement with self-similar core-collapse solutions, and the core collapse times generally agree with expectations from the literature. Also, we observe good total energy conservation, within $\lesssim 0.04\%$ throughout all simulations. We analyze the performance of the code, and demonstrate near-linear scaling of the runtime with the number of processors up to 64 processors for $N=10^{5}$, 128 for $N=10^{6}$ and 256 for $N=10^{7}$. The runtime reaches saturation with the addition of processors beyond these limits, which is a characteristic of the parallel sorting algorithm. The resulting maximum speedups we achieve are approximately 60\texttimes{}, 100\texttimes{}, and 220\texttimes{}, respectively.
|
The dynamical evolution of dense star clusters is a problem of fundamental importance in theoretical astrophysics. Important examples of star clusters include globular clusters, spherical systems containing typically $10^{5}$ - $10^{7}$ stars within radii of just a few parsec, and galactic nuclei, even denser systems with up to $10^{9}$ stars contained in similarly small volumes, and often surrounding a supermassive black hole at the center. Studying their evolution is critical to many key unsolved problems in astrophysics. It connects directly to our understanding of star formation, as massive clusters are thought to be associated with major star formation episodes, tracing the star-formation histories of their host galaxies. Furthermore, globular clusters can trace the evolution of galaxies over a significant cosmological time span, as they are the brightest structures with which one can trace the halo potential out to the largest radii, and they are very old, potentially even predating the formation of their own host galaxies. Unlike stars and planetary nebulae, globular clusters are not simply passive tracers of galaxy kinematics as their internal dynamics are affected by the galactic tidal field. Therefore, their internal properties and correlations with their host galaxies are likely to contain information on the merger history of galaxies and haloes. Dynamical interactions in dense star clusters play a key role in the formation of many of the most interesting and exotic astronomical sources, such as bright X-ray and gamma-ray sources, radio pulsars, and supernovae. The extreme local stellar densities, which can reach $\gtrsim 10^{6}\,{\rm pc^{-3}}$, give rise to complex dynamical processes: resonant stellar encounters, tidal captures, physical collisions, and high-speed ejections (\citealt{2003gmbp.book.....H}). The primary challenge in modeling dense clusters lies in the tight coupling of these processes and their scales as they influence and inform one another both locally, e.g., through close encounters or collisions on scales of $\sim 1-100\,{\rm R_{\odot}}$, or $10^{-8}-10^{-6}$ pc, and globally on the scale of the whole system through long-range, gravitational interactions. Close binary interactions can occur frequently, every $\sim 10^{6}-10^{9}$ yr depending on the cluster density, relative to the global cluster evolution timescale. Furthermore, in the time between close encounters, stellar particles, single and binary, change their physical properties due to their internal nuclear evolution and due to mass and angular momentum transfer or losses. All these changes affect the rates of close encounters and couple to the overall evolution of the cluster. Given these enormous ranges in spatial and temporal scales, simulating dense star clusters with a million stars or more is a formidable computational challenge. A thorough analysis of the scaling of the computational cost of direct $N$-body methods is presented in \citet{1988Natur.336...31H}. Although direct $N$-body methods are free of any approximations in the stellar dynamics, their steep $\propto N^{3}$ scaling has limited simulations to an initial $N\sim10^{5}$ stars (\citealp{2011MNRAS.411.1989Z,2012A&A...538A..19J,2012arXiv1208.4880H}). However, the number of stars in real systems like globular clusters and galactic nuclei can be orders of magnitude larger. Even for globular cluster size systems where the evolution is feasible to calculate with a direct $N$-body code, the total runtime "takes the better half of a year" \citep{2012arXiv1208.4880H} and statistical results have to rely on only a very few models. This is a problem, given the significant inherent stochasticity of these systems, which affects even basic structural parameters \citep[e.g.,][]{2009MNRAS.397L..46H,2010ApJ...708.1598T,2012arXiv1208.4880H}. In order to draw statistically robust conclusions, a much larger number of realizations of massive star clusters has to be calculated, in addition to a wider range of initial conditions. It is clear that these requirements result in prohibitive runtimes for direct $N$-body codes. Monte Carlo methods calculate the dynamical evolution of the cluster in the Fokker-Planck approximation, which applies when the evolution of the cluster is dominated by two-body relaxation, and the relaxation time is much larger than the dynamical time. In practice, further assumptions of spherical symmetry and dynamical equilibrium have to be made. The H\'enon Monte Carlo (MC) technique \citep{1971Ap&SS..14..151H} which is based on orbit averaging, represents a balanced compromise between realism and speed. The MC method allows for a star-by-star realization of the cluster, with its $N$ particles representing the $N$ stars in the cluster. Integration is done on the relaxation timescale, and the total computational cost scales as $N\log N$ \citep{1971Ap&SS..14..151H}. Our work here is based on the H\'enon-type MC cluster evolution code CMC (\textquotedblleft{}Cluster Monte Carlo\textquotedblright{}), developed over many years by \citet{2000ApJ...540..969J,2001ApJ...550..691J,2003ApJ...593..772F,2007ApJ...658.1047F,2010ApJ...719..915C,2012ApJ...750...31U}. CMC includes a detailed treatment of strong binary star interactions and physical stellar collisions \citep{2007ApJ...658.1047F}, as well as an implementation of single and binary star evolution \citep{2010ApJ...719..915C} and the capability of handling the dynamics around a central massive black hole \citep{2012ApJ...750...31U}. In addition to CMC, other MC codes have been developed recently that are based on the same orbit averaging technique. Apart from differences in how the stellar and binary process have been implemented, these codes mainly differ in how particles are advanced in time. The code of \citet{2001A&A...375..711F} uses an individual timestep scheme, where each particle is advanced on its own, local relaxation timescale, while the code of \citet[][with its newest version described in \citet{2011arXiv1112.6246G}]{1998MNRAS.298.1239G} uses a block timestep scheme, where the cluster is divided into several radial zones, and particles are evolved on the average relaxation timescale of the corresponding zone. While they provide better adaptability to density contrasts, individual and block timestep schemes are more difficult to parallelize efficiently in a distributed fashion. A shared timestep scheme, as implemented in CMC, offers a greater degree of inherent parallelism \citep{2001ApJ...550..691J}. A typical simulation starting with $N\sim 10^6$ up to average cluster ages of 12 Gyr using CMC can be run on a modern desktop computer in a reasonable amount of time (days to weeks). However, given the scaling of computational cost, simulations of clusters with $N\gtrsim 10^{7}$ stars, e.g., nuclear star clusters or galactic nuclei, will still take a prohibitive amount of time. Scaling up to even larger number of stars becomes possible only through parallelization. In this paper, we present in detail the latest version of CMC, which is capable of simulating collisional systems of up to $N\sim10^{7}$. In Section \ref{sec:Code-Overview}, we take a look at the components of the serial code and summarize both its numerical and computational aspects. In Section \ref{sec:Parallel-CMC}, we describe the flow of the parallel code, elucidating how we designed each part to achieve optimal performance on distributed parallel architectures. In addition, we describe in the Appendix, an optional CMC feature that accelerates parts of the code using a general purpose Graphics Processing Unit (GPU). We show a comparison of results and analyze the performance of the code in Section \ref{sec:Results}. Conclusions and lines of future work are discussed in Section \ref{sec:Conclusions}.
|
} We presented a new parallel code, CMC (Cluster Monte Carlo), for simulating collisional $N$-body systems with up to $N\sim10^{7}$. In order to maintain a platform-independent implementation, we adopt the Message Passing Interface (MPI) library for communication. The parallelization scheme uses a domain decomposition that guarantees a near-equal distribution of data among processors to provide a good balance of workload among processors, and at the same time minimizes the communication requirements by various modules of the code. Our code is based on the H\'enon Monte Carlo method, with algorithmic modifications including a parallel random number generation scheme, and a parallel sorting algorithm. We presented the first collisional $N$-body simulations of star clusters with $N$ covering three orders of magnitude and reaching up to $N=10^{7}$. The core collapse times obtained in our simulations are in good agreement with previous studies, providing basic validation of our code. We also tested our implementation on 1 to 1024 processors. The code scales linearly up to 64 processors for all cases considered, after which it saturates, which we find to be characteristic of the parallel sorting algorithm. The overall performance of the parallelization is impressive, delivering maximum speedups of up to 220\texttimes{} for $N=10^{7}$. Interesting future lines of work may include reducing the communication overhead by overlapping communication with computation. In addition to the distributed memory parallel version, CMC has also an optional feature that accelerates parts of the algorithm using a general purpose Graphics Processing Unit (GPU), described in the Appendix. An important next step towards reaching even higher $N$ values is the development of a hybrid code which can run on heterogeneous distributed architectures with GPUs. With progress along these lines, we may be able to reach the domain of galactic nuclei for the first time. Many galactic nuclei contain massive, dense star clusters, so-called nuclear star clusters, which are thought to be significantly linked to the evolution of their host galaxies \citep[e.g.,][]{2010IAUS..266...58B}, and their properties might still reflect to some extent the details of the formation of the galaxy \citep{2010ApJ...718..739M}. In addition, with their much larger masses and escape velocities, galactic nuclei are likely to retain many more stellar-mass black holes than globular clusters, and, thus, might significantly contribute to the black hole binary merger rate, as well as to the gravitational wave detection rate of advanced LIGO \citep{2009ApJ...692..917M}. Therefore, the study of galactic nuclei with a fully self-consistent dynamical code such as CMC has the potential to make strong predictions for future gravitational wave detection missions, and might give further insights into the evolution of galaxies.
| 12
| 6
|
1206.5878
|
1206
|
1206.4942_arXiv.txt
|
We propose a mechanism for the observed non-keplerian motion \citep{Giantprops10} of ``propeller'' moons embedded in Saturn's rings. Our mechanism, in which radial variations in surface density -- external to, and unaffected by, the embedded moon -- result in an equilibrium semimajor axis for the moon due to ``Type~I'' angular momentum exchange \citep{Crida10}, provides a good fit to the observations. Future observations should distinguish between our model and others recently proposed.
|
A massive object embedded in a planetary ring creates a propeller-shaped disturbance in the local disk continuum \citep{SS00,SSD02,Seiss05}. Swarms of such ``propellers,'' each with central moonlet radii~$r\sim100$~m, occur in the 3,000-km-wide ``Propeller Belts'' region in the middle part of Saturn's A~ring \citep{Propellers06,Propellers08,Sremcevic07}, though photometric ambiguities make it difficult to precisely ascertain the central moonlets' masses \citep{Anparsgw10}. A second propeller-rich region has been identified in the outermost regions of the A~ring \citep[hereafter T10]{Giantprops10}. Propellers in this region are both larger and rarer than in the Propeller Belts, and several have been tracked for periods of several years. The propeller nicknamed ``Bl\'eriot''\fn{Propellers that have been observed repeatedly are nicknamed to facilitate identification.} is the largest ($r\sim1$~km) and best-tracked ($>100$ detections over 5~years). Its orbit is predominantly keplerian, with longitude residuals less than $\pm0.15^\circ$ (200~km) over a period of nearly 5~yr. However, the residuals are much larger than measurement error and are clearly systematic (T10; data reproduced in \Fig{}~\ref{bleriot_orbit}). If the observed non-keplerian motion is interpreted in terms of changes in the instantaneous semimajor axis $a$, then Bl\'eriot migrated outward from mid-2006 to mid-2007 at a rate of $\dot{a}=+0.11$~km/yr, and inward from late-2007 to early-2009 at a rate of $\dot{a}=-0.04$~km/yr (T10). \begin{figure*}[!t] \begin{center} \includegraphics[width=15cm,keepaspectratio=true]{propeller_orbit3_gpe2-eps-converted-to.pdf} \caption{Observed longitude of the propeller ``Bl\'eriot'' over 4~years, with linear trend $\dot{\lambda}_0=616.7819329^\circ$~day$^{-1}$ subtracted off. Data reproduced from T10, q.v. for more detail. Panel (a) contains all the data, while panels (b), (c), and (d) contain subsets of the data shown in greater detail. The blue line indicates a linear-plus-sinusoidal fit to all the data, while the red lines indicate piecewise quadratic fits corresponding to a constant drift in semimajor axis (T10). The black lines indicate exponential fits as described in Section~\ref{Analytical}. The data from 2005 and 2009 are too sparse to distinguish among fitting functions, so are fit to a simple linear trend. \label{bleriot_orbit}} \end{center} \end{figure*} A linear trend in semimajor axis, that is to say constant $\dot{a}$, corresponds to a quadratic trend in the longitude residual, $\lambda(t)\propto{}t^2$. The observed longitude residuals for Bl\'eriot can be interpreted in terms of a series of piecewise quadratics (red lines in \Fig{}~\ref{bleriot_orbit}) that would imply episodic migration, perhaps caused and punctuated by periodic encounters or collisions \citep[T10]{LS09,Kirsh09}. While plausible, this mechanism needs further development, and this paper will not discuss it further. Other mechanisms for Bl\'eriot's non-keplerian motion have been proposed, and are reviewed below: The observed longitude residuals for Bl\'eriot can be interpreted in terms of a sinusoidal oscillation with period 3.7~yr (blue line in \Fig{}~\ref{bleriot_orbit}). Although the most obvious physical mechanism to produce a sinusoidal longitude residual is resonant interaction with a larger moon exterior to the rings, a mechanism which governs many phenomena in the rings \citep[e.g.,][]{Ringschapter12}, no resonance is known near Bl\'eriot's position that could plausibly cause longitude residuals with the observed large amplitudes. In the ``frog'' mechanism suggested by \citet[hereafter PC10]{PC10} and further developed by \citet{PC12}, the propeller moonlet interacts primarily with the mass at either end of the propeller gap. Modeling those ends as co-orbiting masses, PC10 found that the propeller moonlet plausibly librates with the observed amplitude and period. However, one wonders whether the moon-formed gap responds sluggishly enough to allow the moon to librate within it. Two predictions of this may soon be tested: that Bl\'eriot's 3,000-km-long gap structure (which is only seen clearly in a few images) is stationary with respect to the librations, and that the longitude residual will continue to follow a sinusoidal profile into the future. Other hypotheses for Bl\'eriot's non-keplerian motion, including the one we present below, rely on the concept of ``Type~I migration.'' As classically formulated for protoplanetary disks \citep{Ward86,Ward97,PapaloizouPPV07}, the angular momentum exchange at inner Lindblad resonances between a disk and an embedded mass fails to exactly cancel with that at outer Lindblad resonances, resulting in a differential torque that leads to inward migration of the embedded mass. However, classical Type~I migration depends crucially on the gas component of the disk, which causes Lindblad resonance locations to shift asymmetrically. For the case of planetary rings, which are strictly particulate, \citet[hereafter C10]{Crida10} re-derived the equations for Type~I migration from first principles, using analytical arguments and numerical simulations to trace the angular momentum exchange between streamlines of continuum ring particles and the embedded moon, and deriving a profile of torque per unit disk surface density as a function of the impact parameter $b$ (i.e., the difference in semimajor axis) of ring particle streamlines with respect to the moon (\Fig{}~\ref{cridacalc}). The primary contribution does not depend on resonances, which are more symmetrically placed in a non-gaseous particulate disk, but on intrinsic asymmetries in the single-impulse transfer of angular momentum. \begin{figure*}[!t] \begin{center} \includegraphics[width=8cm,keepaspectratio=true]{cridacalc4-eps-converted-to.pdf} \caption{Angular momentum exchange $\Delta{}J$ between a ring streamline and an embedded moon as a function of the normalized impact parameter $\hat{b}\equiv{}b/r_H$. $\Delta{}J$ is normalized by the moon's specific angular momentum $J_m\equiv{}a_0^2n_0$ and by its Hill radius $r_H$. Data from \citet{Crida10}. \label{cridacalc}} \end{center} \end{figure*} C10 find an intrinsic asymmetric torque, akin to classic Type~I migration, in the case of a particulate disk with uniform surface density; it is always inward (while Bl\'eriot is seen to move both inward and outward) and is one to two orders of magnitude too weak to explain the magnitude of Bl\'eriot's observed non-keplerian motion (C10). However, it is known that Saturn's rings do not have uniform surface density. \citet[hereafter RP10]{RP10} considered stochastic temporal variations of the surface density due to self-gravity wakes in the region of maximum angular momentum transfer, finding that they induce a random walk in the propeller moon's semimajor axis. Under certain realistic conditions, the magnitude of the longitude residual approaches that observed for Bl\'eriot, and since a random walk has no preferred frequency, any frequency might be apparent over a limited period of time. The model of RP10 predicts that the quasi-sinusoidal behavior observed by T10 will not continue, but will instead be seen as part of an overall random progression.
|
} \begin{figure}[!t] \begin{center} \includegraphics[width=13.25cm,keepaspectratio=true]{typeiprops_data-eps-converted-to.pdf} \caption{Normalized brightness ($I/F$) profiles of the A~ring in the vicinity of Bl\'eriot's semimajor axis (vertical dotted line) from the following images: N1540681631, taken on 2006~October~27 (green); N1541717040, taken on 2006~November~8 (cyan); N1560309735, taken on 2007~June~12 (blue); N1595337202, taken on 2008~July~21 (red); and N1654254507, taken on 2010~June~3 (black). Regions of maximum angular momentum exchange are shaded in gray. \label{typeiprops_data}} \end{center} \end{figure} We have shown that a modest permanent radial variation in the surface density, combined with plausible occasional ``kicks'' to the semimajor axis, can lead to non-keplerian motions of an embedded moonlet that are similar to those observed for Bl\'eriot. It finally remains to inquire whether such variations may actually exist in Bl\'eriot's vicinity. This region is characterized by variable structure (\Fig{}~\ref{typeiprops_data}), most likely due to spiral density waves. Our model is viable if, when averaged over orbital timescales (which are short compared to the migration rate), the mean torques due to that structure yield anomalies on either side of Bl\'eriot's position. Optical-depth variations in this part of the A~ring are $\pm0.2$ (M.~Hedman, personal communication, 2011). Using a mass extinction coefficient for the outer-A~ring of 0.01--0.02~cm$^2$~g$^{-1}$ \citep{Colwell09}, this corresponds to surface density variations as high as $\pm10$~g~cm$^{-2}$. The mechanism described herein requires surface density anomalies of only a few g~cm$^{-2}$. Although we cannot prove whether the observed surface density variations are underlain by the much smaller average anomalies necessary to activate our mechanism, nor is such proof likely to be possible without greatly improved observations and/or difficult numerical modeling, our purpose here is only to argue that it is plausible. The model presented in this paper is motivated by the observed orbital evolution of Bl\'eriot and other propellers, which seems to be characterized by periodic ``kicks,'' with slow variations in the semimajor axis (seen as curvature in the longitude residuals) occurring between those kicks. It is in response to these observations that we have constructed a mechanism that is dynamically plausible and consistent with existing data. Our proposed mechanism joins two others recently outlined, and all three mechanisms make contrasting predictions regarding the future behavior of Bl\'eriot's azimuthal residual. While PC10 predict that the quasi-sinusoidal trend discerned in the azimuthal residual to date will continue, and RP10 predict that variations in the azimuthal residual will be seen in the future to result from stochastic variation in the semimajor axis, we predict periodic (less than one per year, on average, from our interpretation of existing data) sudden ``kicks'' to the semimajor axis followed by asymptotic return to the equilibrium value. Future data, beginning in 2013 when \Cassit{} will again spend significant time out of Saturn's equatorial plane with good viewing of the rings, should distinguish among these models. \\ \\ \textbf{Acknowledgements} I thank J.~Burns, M.~Hedman, P.~Nicholson, D.~Hamilton, M.~Evans, A.~Crida, and a reviewer for helpful comments, and I further thank A.~Crida for sharing his calculation results. I acknowledge funding from NASA's Outer Planets Research program (NNX10AP94G). This paper is dedicated to~CGT.
| 12
| 6
|
1206.4942
|
1206
|
1206.1083_arXiv.txt
|
The approximate scale-invariance of correlation functions produced by inflation is due to the dilation isometry of de Sitter space combined with the approximate symmetry of the inflaton dynamics under time translation \cite{Creminelli:2010ba} \be \label{eq:tshift} t\rightarrow \tilde t=t+\mathrm{const} \;. \ee In this paper we want to explore the possibility that this symmetry is promoted to the full time reparametrization invariance \begin{equation} t\rightarrow \tilde t(t)\;. \end{equation} Of course this symmetry can be a good approximation only during inflation while it must be eventually broken, similarly to what happens with the standard symmetry~\eqref{eq:tshift}, at the end of inflation, when reheating takes place. This symmetry has recently been studied in the context of Ho\v rava gravity and its healthy extensions \cite{Horava:2009uw,Blas:2009qj,DP1}. In these references the scalar mode describing the preferred foliation has been dubbed `khronon'. See \cite{Mukohyama:2009gg,Wang:2009azb, Izumi:2010yn,Huang:2012ep} for other possible connections between Ho\v rava gravity and the creation of primordial curvature perturbations. We will see that, once this symmetry is enforced, the inflationary dynamics becomes very constrained and unconventional. In particular three features are worth stressing. \begin{enumerate} \item All correlation functions of $\zeta$ are fixed, at the lowest order in derivatives, by only two coefficients, which can be written in terms of the normalization of the power spectrum and the speed of sound of perturbations. This is in contrast with the general case, where at any order in perturbations one can write new operators. \item During inflation the mode wavefunctions have the same form as in Minkowski. This apparently suggests the lack of a proper production of scalar perturbations. However, as we will argue below, this is not true if one considers the inevitable transition to a phase in which the time-reparametrization symmetry is broken. \item The above feature leaves an interesting signature in the correlation functions of the model. Indeed, the "decaying" mode decays much slower than in the conventional case (as $1/a$ instead of $1/a^3$). This has remarkable consequences for the squeezed limits of correlation functions: the standard single-field theorems hold, but only at first order in the momentum of the long mode. One finds corrections at first order and, in particular, one has a $1/k_L^2$ behaviour of the 3-point function in the squeezed limit. Unfortunately, these effects are very suppressed and totally unobservable. Indeed, the field redefinition symmetry itself is such that a time-dependent background wave, which would violate the consistency relations, can be removed and set to zero. Therefore, these effects are not there in the limit of exact field redefinition symmetry and they will only appear once we consider the small breaking of the symmetry. \end{enumerate} Section \ref{sec:action} describes the construction of the action compatible with the $t\rightarrow \tilde t(t)$ symmetry. The power spectrum is studied in Section \ref{sec:power}, with some details left to the two Appendices. The 3- and 4- point functions are discussed respectively in Section \ref{sec:3pf} and \ref{sec:4pf}, while conclusions are drawn in Section \ref{sec:conclusions}.
|
Conclusions and outlook} Given the simplicity of single-field inflation, it is certainly worthwhile exploring all the possible symmetries that can be imposed on its dynamics and their phenomenological consequences. Here we have studied the implications of imposing an approximate field redefinition symmetry $\phi \to \tilde\phi(\phi)$ on the inflaton. The predictions are very sharp since---after fixing the normalization of the spectrum---all correlation functions depend only on the speed of sound $c_s$ and are somewhat unusual, as a consequence of the slow decay of the decaying mode during inflation. What we have studied represents another de Sitter limit of inflation, as inflation can (but need not) take place with the metric being exactly de Sitter. This parallels the case of ghost inflation \cite{ArkaniHamed:2003uz}, while another example has been studied in \cite{Cheung:2007st}. Like in the case of ghost inflation, the dynamics that may be responsible for modification of gravity in the late Universe, can be applied to inflation. This is not surprising, as models of modification of gravity often involve a scalar which defines a preferred foliation of space-time. And this is exactly what we need for inflation. It is useful to think about this model as another corner of the EFT of inflation \cite{Cheung:2007st}. Starting from a general situation, the limit $\dot H \to 0$ kills the unitary gauge operator $g^{00}$, and therefore the standard spatial kinetic term of the inflaton. This is the limit of ghost inflation \cite{ArkaniHamed:2003uz}, when the spatial kinetic term is given by higher order spatial derivatives ($K^2$ and $K_{\mu\nu} K^{\mu\nu}$), while a standard time kinetic term $\dot\pi^2$ comes from the unitary gauge operator $(g^{00}+1)^2$. The symmetry that we discussed forbids any operator of the form $(g^{00}+1)^n$, so that also the time kinetic term is now given by the higher derivative operator $N^{-2} (\partial_i N)^2$. Of course these are only limiting cases: intermediate regimes in which various operators are relevant may have interesting features. We leave this to future investigations. It is important to stress a relevant drawback of our model, i.e. its spatial non-locality: the Green function of $\pi$ shows instantaneous propagation of the signal as discussed in \cite{Blas:2011ni}. Most likely, this implies that our EFT cannot be embedded in a standard Lorentz invariant UV completion\footnote{We thank D.~Baumann for useful correspondence about this point.}. This is similar to what happens in models of $k$-inflation with superluminal speed of sound $c_s>1$.
| 12
| 6
|
1206.1083
|
|
1206
|
1206.3086_arXiv.txt
|
We apply an extended version of the SU(3) parity model, containing quark degrees of freedom, to study neutron stars. The model successfully reproduces the main thermodynamic features of QCD which allows us to describe the composition of dense matter. Chiral symmetry restoration is realized inside the star and the chiral partners of the baryons appear, their masses becoming degenerate. Furthermore, quark degrees of freedom appear in a transition to a deconfined state. Performing an investigation of the macroscopic properties of neutron stars, we show that observational constraints, like mass and thermal evolution, are satisfied and new predictions can be made.
|
The study of strong interaction physics under extreme conditions is a central topic of nuclear physics with a large number of experimental and theoretical programs focusing on the area. These conditions comprise large temperatures, densities, as well as extreme values of nuclear isospin. In ultra-relativistic heavy-ion collisions a very hot fireball is created in the collision zone and, at high temperatures, hadronic matter is assumed to melt into its constituents, quarks and gluons. The net baryon density in such reactions is determined by the beam energy. At the planned energies in the upcoming FAIR (Facility for Antiproton and Ion Research) at GSI, a hot and relatively dense system will be produced. A central point of these investigations is the understanding of the quark-hadron phase transition. However, while relatively high densities and finite temperatures might be reached in laboratory experiments, the study of neutron stars is essential if one wants to probe the low temperature and high density region of the phase diagram of strongly interacting matter. Observations of neutron star masses (and possibly radii) provide the standard way of constraining the inner composition of these objects. More precisely, recent observations have set new high mass constraints for neutron stars (PSR J1614-2230 \cite{Demorest:2010bx} being the most important), and equations of state aimed to describe compact stars are thus expected to provide objects with high masses \cite{arXiv:1006.5660,arXiv:1107.2497,arXiv:1112.0234,Djapo:2008au,Whittenburyetal.(2012)}. In the case of hybrid stars (neutron stars with a quark core in its center), a quark phase based on a simple non-interacting quark model like the MIT bag model tends to reduce the maximum mass significantly (see the discussion in \cite{arXiv:1011.2233}). A quark phase that includes strong repulsive interactions, however, may have an equation of state quite similar to a purely hadronic one. This prevents the softening of the matter and the drop in maximum mass \cite{nucl-th/0411016,arXiv:1102.2869,arXiv:1108.0559}. It has been found, however, that models with a strong repulsive quark-quark interaction make the description of lattice results at $\mu_B=0$ nearly impossible \cite{Kunihiro:1991qu,Ferroni:2010xf,Steinheimer:2010sp}, since including a repulsive mean field interaction strongly decreases the quark number susceptibility. Therefore, most current successful hybrid star models would fail in this regard, if they were to be applied to the high-temperature, low-density regime. At high temperatures, QCD exhibits a crossover to a deconfined phase and the quarks and gluons become the dominant degrees of freedom \cite{Aoki:2006we}. The temperature at which this transition takes place is estimated to be $T_{dec}\approx 150 - 160$ MeV \cite{Borsanyi:2010cj,Bazavov:2010sb}. A phase transition is also expected to take place at high densities, where baryons start to overlap, and quarks and gluons become the effective degrees of freedom. This indicates that at some point a hadronic model will not be able to appropriately describe the matter present inside a neutron star, and a deconfinement mechanism needs to be introduced. In the following, we will discuss a theoretical approach that is able to describe the conditions found in compact stars as well as those created in heavy-ion collisions. The aim of this approach is to find a unified description for the thermodynamic properties of QCD which can be applied to compact stars and heavy ion collisions at different beam energies while being in accordance with lattice QCD results at vanishing baryon density. An approach of this type is essential if one wants to investigate a phase diagram that has a region of a cross-over transition to quark matter (as it is clearly established by lattice QCD calculations) and a first-order transition at high densities and low temperatures. This is not possible by combining separate models for the hadronic and quark phase. Such a first-order transition at low temperature has also been observed in recent lattice monte-carlo calculations of an effective QCD Lagrangian \cite{philipsen}. In addition, a model that can cover the physics at high temperature as well as at high density can serve as an important tool for the increasingly important studies of black hole formation and black-hole neutron-star mergers, where temperatures up to 90 MeV and high densities might be reached \cite{arXiv:1102.3753,shen}. In a previous paper, an extended quark $\rm SU(3)_f$ parity doublet model was introduced for this purpouse \cite{Steinheimer:2011ea}. In the parity model \cite{Detar:1988kn,Hatsuda:1988mv}, the mass splitting between the nucleons and the respective chiral partners is generated by the spontaneous breaking of chiral symmetry and their coupling to the corresponding order parameter, the scalar (sigma) field. The same applies for the baryon octet in the $\rm SU(3)_f$ case \cite{Nemoto:1998um,Jido:1999hd}. When chiral symmetry is restored and the sigma field vanishes, the chiral partner masses become degenerate. A good description of nuclear saturation properties, as well as neutron star observables can easily be achieved within this formalism \cite{Dexheimer:2007tn,Dexheimer:2008cv}. In Ref. \cite{Sasaki:2010bp} a phase diagram for chiral symmetry restoration was calculated using an $\rm SU(2)_f$ version of the parity model. In addition, the parity model has been shown to describe the lattice results at $\mu_B = 0$ and can be used in dynamical models for relativistic heavy ion collisions \cite{Steinheimer:2009nn,Santini:2011zw,Petersen:2011sb}. We will apply this model to the study of neutron stars. A complete phase diagram for iso-spin symmetric matter, using the extended version of the $ \rm SU(3)_f$ parity model, which contains quark degrees of freedom, has been calculated in \cite{Steinheimer:2011ea}. As expected, at low temperatures the nuclear matter liquid-gas phase transition and the chiral symmetry restoration are of first order. Within this approach the deconfinement phase transition, on the other hand, is a crossover. Furthermore, chiral symmetry restoration and deconfinement do not coincide, and at intermediate densities the matter is chirally symmetric but still confined. Whether such a chirally symmetric hadronic phase can be the $N_c=3$ equivalent of the $N_c=\infty$ quarkyonic phase \cite{McLerran:2007qj} is still subject to debate \cite{Lottini:2011zp,Bonanno:2011yr,Giacosa:2011uk}. In this paper we investigate the properties of electrically-neutral chemically-equilibrated matter in the framework of the model described above. We will investigate the influence of the chiral partners, hyperons and quark matter on the macroscopic properties of a neutron star, such as its gravitational mass, radius, and thermal evolution.
|
We applied for the first time the $\rm SU(3)_f$ version of the parity doublet model to cold, dense, charged neutral and chemically equilibrated matter. Note that this approach is able to successfully describe the low density as well as the high density regime of the QCD phase diagram, as it includes a self-consistent deconfinement transition to quark matter. With these ingredients, we can study for the first time the interplay between the baryon octet, their respective chiral partners and quarks in neutron stars. The chirally symmetric but still confined matter obtained in Ref. \cite{Steinheimer:2011ea} disappears when charge neutrality and chemical equilibrium are taken into account. This is due to the early appearance of the down quark, that compensates the positive charge of the proton. Such a partial appearance of the quarks was already suggested in \cite{Blaschke:2008gd}. The up quark appearance happens, for both cold and warm stars, approximately in the same region as the chiral partners (signaling the chiral symmetry restoration). The zero temperature calculation using the same parametrization for the model as the one calibrated for low and zero chemical potential (Model A) yields stars with masses in agreement with the most massive observed pulsar (1.97~$\pm$ 0.04M$_\odot$ \cite{Demorest:2010bx}). This parametrization, however, leads to nuclear matter with high compressibility. On the other hand, parametrizations leading to more realistic compressibility values (like model B) yield less massive stars. As already shown in Ref. \cite{Dexheimer:2008cv}, corrections that account for the baryonic Dirac sea effect such as the Relativistic Hartree Approximation (RHA) can improve this situation. We note that the reason we did not consider any extra features that could possible improve the situation in this work is because we wanted to study in detail the population distribution in the star taking into account the relation between baryon chiral partners and quarks at high densities, since this had never been performed before. Work along this line is in progress. The inclusion of finite temperature in the calculation allows more massive stars in both of the cases studied, but still does not qualitatively change the situation presented above. We have performed cooling simulations for neutron stars whose microscopic composition is given by our (cold) model. We have found that the presence of the chiral partners affects the thermal evolution, effectively suppressing the hadronic direct Urca process. This, in contrast to other models, yields warmer stars during the neutrino cooling era, and delays the onset of the photon cooling era. Although we cannot effectively test such a prediction at the moment, we will be able to do so, hopefully, in the near future. We also considered the possibility of pairing in the quark phase, where a CFL-like pairing was assumed. The suppression of the quark neutrino emission processes, brought on by pairing, reduces the total emission of neutrinos further, leading to even warmer stars in the neutrino cooling era. We stress that the cooling results presented here should be considered as a first approximation only, since there are many factors that still need to be considered, like the neutrino emissions from the chiral partners, and the effects of stronger quark pairing on the EoS, which warrants future work along those lines. We believe, however, that the cooling investigation put forth in this paper might be a good, qualitative study of the cooling of neutron stars within a $\rm SU(3)_f$ doublet parity model composition, given that the available cooling data can be reproduced. Furthermore, a recent work \cite{Negreiros:2012b} has shown that rotation might have an important effect on the cooling of compact stars, thus we intend to perform 2D simulations of the thermal evolution of stars described by our model.
| 12
| 6
|
1206.3086
|
1206
|
1206.6947_arXiv.txt
|
Elliptical galaxies contain X-ray emitting gas that is subject to continuous ram pressure stripping over timescales comparable to cluster ages. The gas in these galaxies is not in perfect hydrostatic equilibrium. Supernova feedback, stellar winds, or active galactic nuclei (AGN) feedback can significantly perturb the interstellar medium (ISM). Using hydrodynamical simulations, we investigate the effect of subsonic turbulence in the hot ISM on the ram pressure stripping process in early-type galaxies. We find that galaxies with more turbulent ISM produce longer, wider, and more smoothly distributed tails of the stripped ISM than those characterised by weaker ISM turbulence. Our main conclusion is that even very weak internal turbulence, at the level of $\la 15\%$ of the average ISM sound speed, can significantly accelerate the gas removal from galaxies via ram pressure stripping. The magnitude of this effect increases sharply with the strength of turbulence. As most of the gas stripping takes place near the boundary between the ISM and the intracluster medium (ICM), the boost in the ISM stripping rate is due to the ``random walk'' of the ISM from the central regions of the galactic potential well to larger distances, where the ram pressure is able to permanently remove the gas from galaxies. The ICM can be temporarily trapped inside the galactic potential well due to the mixing of the turbulent ISM with the ICM. The galaxies with more turbulent ISM, yet still characterised by very weak turbulence, can hold larger amounts of the ICM. We find that the total gas mass held in galaxies decreases with time slower than the mass of the original ISM, and thus the properties of gas retained inside galaxies, such as metallicity, can be altered by the ICM over time. This effect increases with the strength of the turbulence, and is most significant in the outer regions of galaxies.
|
Ram pressure stripping removes gas from galaxies moving relative to the ICM \citep{1972ApJ...176....1G}. Numerous theoretical studies examined the consequences of this effect for the galaxy and cluster evolution by quantifying the amount of gas loss from galaxies, including star formation in galaxies and their ram-pressure stripping tails, and determining the metal enrichment of the ICM by stripping metal-rich gas from galaxies \citep[e.g.,][]{1994ApJ...437...83B, 2006A&A...452..795D, 2006MNRAS.369.1021M,hester06,2007MNRAS.380.1399R,2007ApJ...671.1434T, 2008MNRAS.383..593M,2008MNRAS.388L..89R,2008ApJ...684L...9T,2008MNRAS.389.1405K, 2009A&A...499...87K,2009A&A...500..693J,2010ApJ...716..810B,2010MNRAS.408.2008T,2011ApJ...729...11K,2011MNRAS.415..257N}. Previous theoretical investigations of ram pressure stripping did not include all non-thermal energy components in the ISM and ICM. In general, non-thermal components include turbulent kinetic energy, magnetic fields, and cosmic-rays. Only few theoretical studies \citep[e.g.,][]{1982MNRAS.198.1007,2003A&A...402..879O, 2006A&A...453..883V,2010NatPh...6..520P} incorporated some of these components. First simulations of ram pressure stripping including the dynamical effects of the magnetic fields were presented in \citet{2012arXiv1203.1343R} for late-type galaxies. We aim to systematically investigate how non-thermal components of the ISM and ICM affect ram pressure stripping in elliptical galaxies. This is our first paper in a series of papers on this subject, and it focuses on the effect of turbulent ISM on the ram pressure stripping rates, morphologies of the stripping tails, and mixing between the ISM and ICM. Although observational constraints on the turbulence properties of the hot ISM in early-type galaxies are still uncertain, there is little doubt that the hot ISM is characterised by weak turbulence and randomly oriented weak magnetic fields \citep{2003ARA&A..41..191M,2004ARA&A..42..211E,2010MNRAS.402L..11S,2012arXiv1205.0256H}. Recent X-ray observations began to place meaningful constraints on the magnitude of turbulent motions in the hot gas of massive early-type galaxies, proving that the turbulence is subsonic \citep[e.g.,][]{2008MNRAS.388.1062C,2010MNRAS.404.1165C, 2010MNRAS.406..354O,2011MNRAS.410.1797S,2012A&A...539A..34D,2012ApJ...747...32B}. Stellar winds, supernovae, and active galactic nuclei are considered to be the main energy sources for these turbulent motions \citep{1996MNRAS.279..229M,2009ApJ...699..923B,2011ApJ...728..162D}. We study the morphology of the ram pressure stripping tails. Sharp edges characteristic of ram pressure stripping have been detected in X-ray maps of a galaxy falling into the Fornax cluster \citep{2005ApJ...621..663M}, and X-ray tails are sometimes observed in ellipticals undergoing ram pressure stripping \citep[][]{2008ApJ...688..208R, 2008ApJ...688..931K, 2006ApJ...644..155M}. When the strength of stripping is not significant, or the duration of the process is short, galaxies are likely to show only somewhat elongated gas distributions instead of long tails \citep[e.g.,][]{2010MNRAS.405.1624M}. In addition to the morphology of the gas distribution, ram pressure stripping also can be probed by tracking how well the ISM and ICM are mixed together for different ISM stripping rates \citep[e.g.,][]{2005A&A...435L..25S}. Most previous simulations focused on how material is expelled from galaxies into the ICM. Here we also examine how much mass can be mixed into galaxies from the ICM due to the random motions of the ISM in ram pressure stripping. A direct consequence of this mixing can be a significant change of the metallicity in the galactic gas. The organisation of the paper is as follows. We describe the simulation setup in Section 2. In Section 3, we present results of our simulations, emphasising the differences in the impact of various strengths of turbulent motions on the ram pressure stripping rates and tail morphologies. Finally, we present conclusions and discussion in Section 4.
|
We show that the continuous supply of small to moderate amount of turbulent kinetic energy to the ISM enhances the ISM mass loss rate in elliptical galaxies experiencing ram pressure stripping, and increases the penetration of the ICM into the galaxies (see Figure \ref{fig:mass_ISM_ICM}). The spatial distribution of the stripped ISM can be wider and more extended along the direction of galaxy motion (see Figures \ref{fig:3D}, \ref{fig:color_global}, and \ref{fig:particle_global}), when AGN feedback and/or stellar processes such as star formation are present. Our results imply that early-type galaxies characterised by the turbulent ISM should efficiently disperse their ISM throughout galaxy clusters. The origin of the stripped ISM in the tails shows that the ram pressure stripping with the turbulent motions in the ISM boosts the mixing between the central region and the outer region of the galaxy. This implies that the distributions of gas properties in the tails can be used to infer the distribution of the intrinsic ISM properties such as gas metallicity inside galaxies. Since the distant part of the tail in Run 5 is more mixed with the central gas inside the galaxy than in Run 0, we expect that the properties of the stripped ISM should show weaker gradients of the gas properties along the tail in Run 5 than in Run 0. For example, there might be a gradient in metallicity distribution along the ram pressure stripping tail. In general, the ISM in the central regions of early-type galaxies is more metal-rich than in the outer regions \citep[e.g.,][]{2011MNRAS.418.2744M, 2011ApJ...729...53H}. Therefore, very low levels of the ISM turbulence in early-type galaxies make the stripping tail have only low metallicity ISM along the tail, contributing negligibly to the ICM metal enrichment \citep[see,][for discussion of the ICM enrichment efficiency]{2008SSRv..134..363S,2008ApJ...688..931K}. However, we note that this depends on the initial metallicity distribution in the galaxy. If a galaxy has a shallow metallicity gradient before experiencing ram pressure stripping, it can flatten the metallicity distribution along the tail and lead to more significant ICM enrichment even when the strength of the ISM turbulence is weaker. The evolution of the mass inside the galaxy implies that a significant fraction of the gas mass measured in observations can be explained by the ICM gas that got temporarily incorporated into the ISM. As Figure \ref{fig:ICM_over_ISM} shows, galaxies with the strong turbulent motions in the ISM easily blend the inflowing ICM with the ISM. Therefore, the properties of the hot X-ray emitting ISM in the galaxy experiencing ram pressure stripping might have been altered by the inflowing ICM, in particular, in the outer regions. For example, the metallicity of the ICM in low-redshift galaxy clusters is about 0.5 $Z_{\odot}$ \citep{2009ApJ...698..317A} and can be much lower than that of the ISM \citep{2006ApJ...639..136H,2009ApJ...696.2252J}, Thus, mixing of the ICM with galactic gas can alter the metallicity of the ISM and the metallicity gradient inside cluster galaxies. This contamination can be particularly significant in the outer regions of galaxies when the ISM is turbulent. Interestingly, if the galaxy has an initially flat metallicity profile at the level of 2 $Z_{\odot}$, and if the metallicity of the ICM is about 0.5$Z_{\odot}$, then the ram pressure stripping will lower the mass-weighted metallicity to around 0.9 and 1.9 $Z_{\odot}$ in 50 $\leq r <$ 100 kpc and $r <$ 20 kpc, respectively (with the mass ratios shown in Figure \ref{fig:ICM_over_ISM} for Run 5 at 6 Gyr) This specific case illustrates how ram pressure stripping in the presence of turbulent ISM can steepen ISM metallicity profiles. This steepening effect is expected to be more pronounced as the strength of ISM turbulence increases. The models presented here allow one to study the effects of turbulence on the ram pressure stripping process via a conceptually simple approach. The advantage of this approach lies in providing a clear intuitive picture of how the turbulent ISM affects the gas stripping. These models form a framework for future studies that will relax some of the assumptions made in the present work. Our current simulations do not include a few important physical processes that are required to make detailed observational predictions for the ram pressure stripping process. First, we do not include radiative cooling processes \citep[see,][for a review]{2008SSRv..134..155K}, which will lead to the formation of dense cold gas clouds \citep[e.g.,][]{2007ApJ...671..190S,2010ApJ...717..147S,2010ApJ...722..412Y}. Second, self-gravity of the gas is not included. Self-gravity can alter the evolution of the stripped ISM by accelerating the collapse of these dense cold gas clouds. Third, the spatial resolution of our simulations is not high enough to fully cover an extremely broad inertial range of the turbulent ISM \citep[see][for a review]{1998AnRFM..30..539M,2011RPPh...74d6901B}. Fourth, we have neglected magnetic fields, which may affect the efficiency of mixing of the ISM and ICM, suppress viscosity and thermal conduction between the stripping tail of the cold gas and the hot ICM, and introduce non-trivial dynamical effects. Finally, the energy sources of the turbulence in our simulations are not directly controlled by the relevant astrophysical processes such as star formation and AGN. Continuous mass loss by ram pressure stripping can affect star formation and AGN feedback \citep{1992MNRAS.255..346B,1999MNRAS.309..161M,2008A&A...481..337K,2012ApJ...745...13S}, increasing or decreasing energy injected to the turbulent ISM. Our main concern is the fact that our model currently does not take into account a possible coupling between the efficiency of stirring of the gas by star formation and AGN and the efficiency of stripping. For example, it is conceivable that enhanced stellar or AGN feedback could increase the level of turbulence, accelerate the mass removal from the galaxy, and thus reduce the fuel supply for these feedback processes and the efficiency of the ram pressure stripping process. Consequently, less gas would be available to fuel AGN and star formation, and the stirring efficiency would slow down. Our model currently does not incorporate such mechanism. However, we show that the efficiency of ram pressure stripping depends sensitively on the duration of stirring, and our models for the continuous (Case A) and initial (Case B) stirring likely bracket the range of possibilities. In future work, we will relax some of the assumptions and simplifications made here. The second paper in this series will investigate the effect of weakly magnetised turbulent ISM in elliptical galaxies on the ram pressure stripping process.
| 12
| 6
|
1206.6947
|
1206
|
1206.4035_arXiv.txt
|
The time domain has been identified as one of the most important areas of astronomical research for the next decade. The Virtual Observatory is in the vanguard with dedicated tools and services that enable and facilitate the discovery, dissemination and analysis of time domain data. These range in scope from rapid notifications of time-critical astronomical transients to annotating long-term variables with the latest modelling results. In this paper, we will review the prior art in these areas and focus on the capabilities that the VAO is bringing to bear in support of time domain science. In particular, we will focus on the issues involved with the heterogeneous collections of (ancilllary) data associated with astronomical transients, and the time series characterization and classification tools required by the next generation of sky surveys, such as LSST and SKA.
|
\label{intro} The time domain is the emerging field of astronomical research, as recognized in the 2010 National Research Councils Decadal Survey of Astronomy and Astrophysics \cite{decadal}. Planned facilities for the next decade and beyond, such as the Large Synoptic Survey Telescope (LSST)\footnote{http://www.lsst.org} and the Square Kilometer Array\footnote{http://www.skatelescope.org} (SKA), will revolutionize our understanding of the universe with nightly searches of large swathes of sky for changing objects and networks of robotic telescopes ready to follow up in greater detail selected interesting sources. This will impact essentially every area of astronomy, from the Solar System to cosmology, and from stellar evolution to extreme relativistic phenomena\cite{ref8}, making it a very rich area for scientific exploration and discovery. Moreover, many interesting phenomena, e.g., supernovae and other types of cosmic explosions, can only be studied in the time domain. These new surveys build on a legacy of over fifty years of experience of sky surveys, first with photographic plates and then, more recently, digital detectors (see Ref.~\citenum{djorgovski} for a recent review). The rise of information technology has driven an exponential growth of data volumes (and, equally importantly, data complexity and data quality) following Moore's law, e.g., DPOSS \cite{ref3} to 2MASS \cite{ref4} to SDSS\cite{ref2}, and many digital sky surveys that followed. To cope with such (necessarily distributed) giga- and terascale data collections, the community developed the concept of the {\em Virtual Observatory} (VO), which provides the wherewithal to aggregate and analyze disparate data sets, opening up new avenues of scientific research based on data discovery and fusion\cite{vo1, vo2}. The Virtual Astronomical Observatory\cite{vao} (VAO) is the US national VO project and provides the components, libraries, and templates that allow national facilities, major projects, and end-users to craft their own VO-enabled applications for seamless data access and integration, especially in support of data intensive research. The time domain has been identified early as a prime arena for VO applications\cite{vo3}. It adds a new dimension to data discovery and federation with (near) real-time massive data streams - for example, Palomar-Quest\cite{pq}, Catalina Real-time Transient Survey\cite{crts} (CRTS), Palomar Transient Factory\cite{ptf} (PTF) and Panoramic Survey Telescope And Rapid Response System\cite{ps1} (Pan-STARRS; PS1) and LSST to come - replacing static data sets; in a way, we've moved from panoramic digital photography of the sky to panoramic digital cinematography. Since many of the observed phenomena in this domain are short-lived, and since the scientific returns depend strongly not only on their detection, but also on the timely and well-chosen follow-up observations, there is a need to fully process the data as they stream from the telescopes, compare it with the previous images of the same parts of the sky, automatically and reliably detect any changes, and classify and prioritize the detected events for the rapid follow-up observations. This poses significant new technological challenges for the VO and its infrastructure. Analogous situations may also be found in many other areas, where the data come continuously from some instruments or sensor networks, and where anomalous or specifically targeted events have to be found and responded to in a rapid fashion. The VO has evolved a two-track approach to the time domain: one deals specifically with the mechanics of reporting transient celestial events (VOEvent) in a timely fashion and the associated infrastructure to publish, disseminate and archive them. The other deals with the more general issues of time series data, such as how to describe, represent and access them in a way to ensure interoperability between different data archives. In this paper, we will review the specific issues that are associated with the latter and how the VO, and, specifically, the VAO, which is leading the time domain effort, is meeting them. This draws in associated but separate work on source characterization and classification that is an essential part of a time domain system. Details of the transient approach are presented in a complementary paper in these proceedings \cite{skyalert-spie}, although we will discuss certain common issues. Nevertheless, whichever approach is being discussed, operational concerns are an important consideration and we focus particularly on those related to questions of scalability and managing large collections of heterogeneous data.
|
The emerging field of time domain astronomy requires tools and infrastructure to support a distributed network of (massive) real-time data streams, data archives, and analysis services. The VAO is developing an interoperable framework to connect partner providers of both data and analysis resources, and expose them as an integrated whole for wider community use. A recent community-wide call for collaborative proposals by the VAO\cite{vao} has identified two time domain projects which it is now advising. One is concerned with access to data related to the Variable and Slow Transient (VAST) Survey Science Program of the Australian Square Kilometre Array Pathfinder (PI: T. Murphy) and the other involves access to the databases of the American Association of Variable Star Observers (AAVSO, PI: M. Templeton). Such collaborations, combining domain expertise in data technologies and the relevant science areas, illustrate the potential of an informatics-based approach to data-intensive science.
| 12
| 6
|
1206.4035
|
1206
|
1206.6801_arXiv.txt
|
We discuss the potential of the eROSITA telescope on board the \emph{Spectrum-X-Gamma} observatory to detect gamma-ray burst (GRB) X-ray afterglows during its 4-year all-sky survey. The expected rate of afterglows associated with long-duration GRBs without any information on the bursts proper that can be identified by a characteristic power-law light curve in the eROSITA data is 4--8~events per year. An additional small number, $\lesssim 2$~per year, of afterglows may be associated with short GRBs, ultra hard (GeV) GRBs and X-ray flashes. eROSITA can thus provide the first unbiased (unaffected by GRB triggering) sample of $\lesssim 40$ X-ray afterglows, which can be used for statistical studies of GRB afterglows and for constraining the shape of the GRB $\log N$--$\log S$ distribution at its low-fluence end. The total number of afterglows detected by eROSITA may be yet higher due to orphan afterglows and failed GRBs. The actual detection rate could thus provide interesting constraints on the properties of relativistic jets associated with collapse of massive stars. Finally, eROSITA can provide accurate ($\lesssim 30''$) coordinates of newly discovered afterglows within a day after the event, early enough for scheduling further follow-up observations.
|
The main objective of the \emph{Spectrum-Roentgen-Gamma} (\emph{SRG}) observatory is to perform a sensitive all-sky survey in the 0.3--12 keV energy band with the eROSITA\footnote{Extended ROentgen Survey with an Imaging Telescope Array} \citep{Predehl11} and ART-XC\footnote{Astronomical Roentgen Telescope -- X-ray Concentrator} \citep{Pavlinsky11} telescopes. The survey (see \S\ref{s:tasks} for details) will last 4 years and consist of 8 repeated complete scans of the sky. The telescopes will be scanning the sky in great circles as a result of the spacecraft's rotation with a period of 4~hours around its axis pointed at the Sun. This observational strategy provides the possibility of studying variable and transient X-ray sources on three characteristic time-scales corresponding to 1) the duration of a single scan of a point source, $\lesssim 40$~s, 2) the duration of a single observation of a source, $\gtrsim 1$~day (consisting of $\gtrsim 6$ consecutive rotation cycles, depending on the ecliptic latitude), and 3) the duration of a single all-sky scan (6~months). Information on temporal behaviour can greatly assist in identifying the types of X-ray sources discovered during the eROSITA all-sky survey. In particular, the X-ray afterglows of cosmic gamma-ray bursts (GRBs) form a class of bright transient sources that usually demonstrate power-law decay during the first hours and days after the burst (see \citealt{gehrels09} for a review). Therefore, GRB afterglows can manifest themselves by a distinct variability pattern on the time-scale of several successive eROSITA scans, which in principle makes it possible to identify such events by analysing the X-ray light curves of sources detected during the all-sky survey. The purpose of this paper is to estimate the detection rates of GRB X-ray afterglows during the eROSITA all-sky survey. Previously, \cite{Greiner00} carried out a search for afterglows of untriggered GRBs in the \textit{ROSAT} all-sky survey (RASS) data and found 23 afterglow candidates. However, a closer examination indicated that at least half and perhaps the majority of these events were flares of late-type stars. Taking into account eROSITA's better (by a factor of $\sim 4$ in the 0.5--2~keV energy band) sensitivity and a factor of 2 larger sky coverage (survey duration times field of view area) of the planned survey, we can expect a significantly larger number of detected GRB afterglows. We note in passing that in addition to afterglows, the \emph{SRG} all-sky survey may also detect a significant number of GRBs themselves, depending on the unknown shape of the $\log N$--$\log S$ distribution of GRBs at very low fluences. This topic will be addressed elsewhere. The main motivation for such a study is that eROSITA can provide the \emph{first unbiased sample of X-ray afterglows}. The problem with all existing samples of afterglows is that they are based on trigerred GRBs and thus determined by the energy range, sensitivity and strategy of the particular GRB experiments. As we discuss in this paper, eROSITA may find a significant number of afterglows associated with GRBs falling near or below the detection threshold and/or having the spectral maximum outside the energy range of existing GRB monitors. This will make it possible to construct an unbiased distribution of X-ray afterglow fluxes and will also provide valuable constraints on the shape of the GRB $\log N$--$\log S$ distribution at low prompt emission fluences. In addition, eROSITA may find a significant number of 'orphan' afterglows, i.e. events without preceding prompt high-energy emission (e.g. \citealt{Rossi2002,Nakar2003}), and afterglows associated with 'failed GRBs' (e.g. \citealt{Huang2002}), which will provide interesting constraints on the properties of relativistic jets produced during the collapse of massive stars \citep{MW1999}.
|
The results of this study imply that it should be possible to detect and identify, by the shape of the light curve, 4--8 X-ray afterglows associated with classical long GRBs per year in the eROSITA all-sky survey data (see Table~\ref{sum}). The exact number will depend on the shape of the $\log N$--$\log S$ distribution of GRBs at low fluences (near the effective threshold of \emph{CGRO}/BATSE and \emph{Swift}/BAT. In addition, eROSITA is expected to find a small number of afterglows associated with other classes of GRBs such short bursts, GeV bursts and X-ray flashes (Table~\ref{typesum}). Thus, by the end of the 4-year survey, a sample of at least 20--40 X-ray afterglows can be accumulated. This sample, although smaller than the already existing samples of afterglows, will nevertheless be interesting for systematic studies of GRBs and their afterglows because of its unbiased nature. In particular, it can be used to construct an unbiased distribution of X-ray afterglow fluxes and to obtain constraints on the shape of the $\log N$--$\log S$ distribution of GRBs. The total number of afterglows detected by eROSITA may prove higher, perhaps by a factor of 2 or more, due to orphan afterglows and failed GRBs. The actual detection rate will thus provide interesting constraints on the properties of relativistic jets associated with the collapse of massive stars. The proposed algorithm for searching afterglows of non-triggered GRBs (Task~1) in the eROSITA data is based on checking if the light curve of a given source resembles a power-law decline. Although such a procedure may erroneously identify a large number of other types of variable X-ray sources (e.g. AGN and stellar flares) as GRB afterglows, most of such contaminants should be easily revealed through cross-checking of successive eROSITA all-sky scans and cross-correlation with large-area optical and infrared source catalogues. We have also discussed the possibility of using the eROSITA data for searching for afterglows of triggered (by any GRB monitors) GRBs. Since the coordinates of the triggered bursts will be known, eROSITA just needs to detect a few photons from the afterglow in one $\sim 40$~s scan. As a result, the total number of such events can be large, $\sim 20-60$~per year depending on the $\log N$-$\log S$ function (see Table~\ref{sum}), provided that at the time of the \emph{SRG} mission there are GRB monitors covering most of the sky at any given time. X-ray afterglows detected in this way can be interesting for statistical studies addressing the same scientific problems as discussed above in relation to the search for afterglows of non-triggered GRBs. Finally, although \emph{SRG} data transfer is planned to occur only once per day, accurate ($\lesssim 30''$) coordinates provided by eROSITA within $\sim 1$~day after the event on afterglows of GRBs and orphan afterglows can be valuable for scheduling further follow-up observations.
| 12
| 6
|
1206.6801
|
1206
|
1206.2339_arXiv.txt
|
We consider the motion of planar phase-transition fronts in first-order phase transitions of the Universe. We find the steady state wall velocity as a function of a friction coefficient and thermodynamical parameters, taking into account the different hydrodynamic modes of propagation. We obtain analytical approximations for the velocity by using the thin wall approximation and the bag equation of state. We compare our results to those of numerical calculations and discuss the range of validity of the approximations. We analyze the structure of the stationary solutions. Multiple solutions may exist for a given set of parameters, even after discarding non-physical ones. We discuss which of these will be realized in the phase transition as the stationary wall velocity. Finally, we discuss on the saturation of the friction at ultra-relativistic velocities and the existence of runaway solutions.
|
In a first-order cosmological phase transition, bubbles nucleate and expand, converting the high-temperature phase into the low-temperature one (see, e.g., \cite{gw81,ah92,m00}). As bubbles expand, latent heat is released at their boundaries. This energy raises the temperature and causes bulk motions of the plasma. The perturbations caused in the cosmic fluid by the nucleation and expansion of bubbles generate a departure from thermal equilibrium. This may give rise to a number of cosmic relics, such as a baryon number asymmetry \cite{ckn93}, baryon inhomogeneities \cite{w84}, magnetic fields \cite{gr01}, topological defects \cite{vs94}, or gravitational waves \cite{gw,lms12}. In general, the system can be described by a relativistic fluid and a scalar field $\phi$ at finite temperature $T$ \cite{ikkl94,kl95,kl96}. The latter may be a Higgs field and acts as an order parameter. At high temperatures the free energy $\mathcal{F}(\phi,T)$ has a minimum $\phi_+(T)$ (in general, $\phi_+\equiv 0$) and at low temperatures a different minimum $\phi_-(T)$. In a first-order phase transition, the two minima coexist in certain range of temperatures, separated by a barrier. In the high-$T$ phase, the free energy density is given by $ \mathcal{F}_{+}(T)= \mathcal{F}(\phi_+(T),T)$, whereas in the low-$T$ phase, it is given by $\mathcal{F}_{-}(T)= \mathcal{F}(\phi _{-}(T),T)$. The critical temperature $T_c$ is that for which $\mathcal{F}_{+}(T_{c})=\mathcal{F}_{-}(T_{c})$. The phase transition occurs when the temperature of the Universe reaches $T_c$. At $T=T_c$, though, the nucleation rate vanishes, and bubbles effectively begin to nucleate at some temperature $T_n$ below $T_c$ \cite{ah92,c77}. The nucleated bubbles expand due to the pressure difference between the two phases. In general, the bubble walls reach a terminal velocity $v_w$ due to the friction with the surrounding plasma. Recently, the hydrodynamics of the moving walls has received much attention (see, e.g., \cite{ms09,bm09,ekns10,lm11,kn11}) due to the interest in performing thorough calculations of the wall velocity and the energy injected into bulk motions of the fluid. These quantities are relevant for the generation of baryons and gravitational waves. In Ref. \cite{bm09}, the ultra-relativistic velocity regime was considered, and it was shown that a state of continuous acceleration of the bubble wall is possible. Such ``runaway'' solutions may play an important role in the generation of gravitational waves. A realistic evaluation of the cosmological consequences of a phase transition requires considering the dynamics as completely as possible. Following the development of a phase transition involves the calculation of several temperature-dependent quantities such as, e.g., the pressure of the two phases and the bubble nucleation rate. During the phase transition the temperature varies in time and space due to the adiabatic cooling of the Universe and the release of latent heat. As a consequence, one has to deal with a set of integro-differential equations. Some of the involved variables are very sensitive to approximations (for instance, the nucleation rate). In order to avoid large errors, it is convenient to resort to nontrivial numerical calculations for these quantities. On the other hand, finding analytical approximations for other variables (e.g., the bubble wall velocity) provides a way of reducing the computation time. Widely used simplifications include the thin wall approximation and the bag equation of state \cite{ikkl94,kl95,kl96,gkkm84}. Even with these approximations, it is not always possible to obtain analytical results. A considerable simplification is achieved by considering planar walls. Analytic results for the planar case were found recently in Ref. \cite{ms09} for the wall velocity and in Ref. \cite{lm11} for the energy injected into the fluid. It is important to note that considering spherical bubbles is not necessarily a better approximation than considering planar walls. Although the spherical symmetry is a good approximation for the initial stages of bubble growth, some cosmologically interesting outcomes of the phase transition are produced when bubbles collide and lose the spherical symmetry. Moreover, losing the spherical symmetry is a requirement, e.g., for the generation of turbulence or of gravitational waves. As an explicit example, the ``envelope approximation'' for the generation of gravitational waves in bubble collisions neglects the overlap regions of colliding bubbles and follows only the evolution of the uncollided bubble walls. For such a calculation, the approximation of treating the walls as planar is, in principle, as good as considering spherical bubbles (but less complicated). In general, one does not expect important differences (see Ref. \cite{lm11} for a comparison of different wall geometries). In this paper we investigate the propagation of a planar phase transition front in the plasma. We aim at finding analytical approximations for the stationary velocity of the bubble wall as a function of the friction and the thermodynamical parameters. The present work is a continuation of the investigations of Ref. \cite{ms09}, where we considered planar walls propagating as weak deflagrations or weak detonations. Here we include into consideration the case of supersonic Jouguet deflagrations \cite{kl95} and the possibility that the walls run away (see Ref. \cite{ekns10} for a recent study for spherical-symmetry walls). We also discuss here whether the different solutions are physical or not, and which of them will be realized as final stationary states during the phase transition. We discuss on the validity of the analytical approximations. The approximations are better for weaker solutions than for those close to the Jouguet points. For comparison, we consider some cases previously studied with numerical calculations \cite{ikkl94,kl95,kl96}. The paper is organized as follows. In section \ref{hydro} we consider the equations for the profiles of the fluid and the bubble wall, including a phenomenological friction term. We study the thin wall limit. In section \ref{bag} we use the bag equation of state to obtain a set of analytic equations for the wall velocity. In sections \ref{test} and \ref{result} we present our results for the stationary motion and compare them with those of the numerical works of Refs. \cite{ikkl94,kl95,kl96}. In section \ref{test} we discuss the range of validity of the analytical approximations and in section \ref{result} we analyze the dependence of the wall velocity on the thermodynamic parameters and the friction. In section \ref{runaway} we consider a different phenomenological friction term, which takes into account the fact that, in some models, the friction force approaches a constant in the ultra-relativistic regime. Finally, we conclude in section \ref{conclu}.
|
\label{conclu} We have investigated the steady state motion of phase-transition fronts in a cosmological first-order phase transition. Our main goal was to find analytical approximations for the wall velocity, taking into account the different possibilities for the hydrodynamical modes and fluid profiles. Therefore, we have considered the case of planar walls, which allow to obtain analytical approximations. In Ref. \cite{ms09} we considered the cases of weak deflagrations preceded by a shock front and weak detonations followed by a rarefaction wave. Here, we have studied also the case of Jouguet deflagrations which have both shock and rarefaction waves and move supersonically. We have considered two different phenomenological models for the friction. One of them grows linearly with the relativistic velocity $\gamma v$ \cite{ikkl94}, and the other saturates for large $\gamma v$ \cite{ekns10}. The latter reproduces the behavior of the friction force in the ultra-relativistic limit and leads to runaway solutions \cite{bm09}. Our main result is a set of algebraic equations which allow to obtain, from the thermodynamic parameters and the friction coefficient, the value of the wall velocity which will be realized as the final stationary state. These analytical results rely on several approximations, such as the use of the bag equation of state and the thin wall approximation. Implementing the latter in the equation for the friction requires some ansatz for the variation of the entropy density inside the wall and also for that of the fluid velocity. The approximation for the entropy density seems to be the roughest one, as we do not obtain the curve of zero entropy production in the limit of vanishing friction. By comparing with numerical lattice calculations \cite{ikkl94,kl95,kl96}, we have checked that the strongest departure from the exact solution occurs for the strongest physical solutions, i.e., those around the Jouguet points (either for detonations or deflagrations). For a friction of the form $\gamma v$, our results are in good quantitative agreement with the cases of planar walls considered in Refs. \cite{ikkl94,kl95} and in good qualitative agreement with those of spherical bubbles considered in Refs. \cite{kl95,kl96}. For a friction which saturates for large $\gamma$, the wall velocity shows essentially the same behavior as in Ref. \cite{ekns10}, which considered spherical bubbles. We remark that this latter approximation, although reproducing the correct behaviors for $v\to 0$ and for $v\to 1$, may still be too simplistic for intermediate velocities. In particular, the friction saturates to a constant value for relatively small velocities, i.e., for values of the gamma factor which are much smaller than those which justify the approximations that lead to the runaway solution \cite{bm09}. As a consequence, the region of parameter space in which the bubble wall runs away may be largely overestimated. The problem of determining the behavior of the friction at intermediate velocities is a difficult one and deserves further investigation, since it has important implications for cosmology.
| 12
| 6
|
1206.2339
|
1206
|
1206.5519_arXiv.txt
|
We present parallax and proper motion measurements, near-infrared spectra, and WISE photometry for the low surface gravity L5$\gamma$ dwarf 2MASSJ035523.37+113343.7 (2M0355). We use these data to evaluate photometric, spectral, and kinematic signatures of youth as 2M0355 is the reddest isolated L dwarf yet classified. We confirm its low-gravity spectral morphology and find a strong resemblance to the sharp triangular shaped $H$-band spectrum of the $\sim$10~Myr planetary-mass object 2M1207b. We find that 2M0355 is underluminous compared to a normal field L5 dwarf in the optical and MKO $J,H$, and $K$ bands and transitions to being overluminous from 3-12 $\mu$m, indicating that enhanced photospheric dust shifts flux to longer wavelengths for young, low-gravity objects, creating a red spectral energy distribution. Investigating the near-infrared color magnitude diagram for brown dwarfs confirms that 2M0355 is redder and underluminous compared to the known brown dwarf population, similar to the peculiarities of directly imaged exoplanets 2M1207b and HR8799bcd. We calculate UVW space velocities and find that the motion of 2M0355 is consistent with young disk objects ($<$ 2-3 Gyr) and it shows a high likelihood of membership in the AB Doradus association.
|
With masses intermediate between stars and planets (i.e., below the hydrogen burning and above the deuterium burning mass limit), brown dwarfs provide a natural link between stellar astrophysics and the planetary science of gas-giants (\citealt{Saumon1996}; \citealt{Chabrier1997}). Studies of the population have informed our understanding of low-mass star formation as well as the physical and chemical composition of low-temperature photospheres (e.g. \citealt{Burrows01,Burrows97}; \citealt{Chabrier00}). With an increasing number of brown dwarf discoveries, the diversity of the population in age, atmospheric properties, and chemical composition is becoming apparent. Brown dwarfs are classified using red optical or near-infrared spectra and show characteristics which distinguish them as L (T$_{eff}\sim$2200 - 1300K) or T/Y (T$_{eff}<$1300) dwarfs (\citealt{Kirkpatrick99}; \citealt{Burgasser02}; \citealt{Cushing11}). The majority of spectrally classified field brown dwarfs within the literature are nearby isolated L dwarfs. Among the $\sim$1000 objects spanning this temperature regime, a significant portion exhibit near-infrared colors, spectral energy distributions (SEDs), and kinematics consistent with a field age population (e.g., \citealt{Kirkpatrick00}; \citealt{Knapp04}; \citealt{Cruz07}; \citealt{Chiu06}; \citealt{Faherty09}; \citealt{Schmidt10}). However there are subsets exhibiting strong deviations in observational properties from the general population including low-metallicity subdwarfs, low surface gravity objects, and potentially cloudy/cloudless L dwarfs (\citealt{Burgasser03,Burgasser04,Burgasser07}; \citealt{Looper08}; \citealt{Cruz09}; \citealt{Cushing09}; \citealt{Kirkpatrick10}; \citealt{Rice10}; \citealt{Radigan12}). The most relevant sub-population to giant exoplanet studies are young (i.e., low surface gravity) isolated L dwarfs. The archetypal low surface gravity L dwarf, 2MASSJ01415823$-$4633574 (2M0141), was discovered by \citet{Kirkpatrick06}. Its optical spectrum exhibits strong bands of VO but abnormally weak TiO, K, and Na absorption. In the near-infrared, its red $J-K_{s}$ color (2MASS $J-K_{s}$=1.73) and triangular $H$-band spectral morphology distinguish it from field L dwarfs (\citealt{Kirkpatrick10}; \citealt{Patience12}). It is suspected to be a member of the $\beta$ Pictoris or Tucana-Horologium association, although the precise kinematics required to confirm association have not yet been determined (\citealt{Kirkpatrick10}). After the discovery and characterization of 2M0141, additional isolated L dwarfs sharing similar photometric and spectral peculiarities attributed to a low surface gravity were reported (e.g. \citealt{Reid08}; \citealt{Cruz09}; \citealt{Kirkpatrick10}). While the ages of these seemingly young L dwarfs remain largely unconstrained, there are kinematic and spatial indications that they represent the lowest mass members of nearby moving groups such as AB Doradus, $\beta$ Pictoris, Tucana-Horalogium (\citealt{Cruz09}; \citealt{Kirkpatrick10}). \citet{Cruz09} point out that the majority of objects defining the population of the lowest surface gravity L dwarfs show spectral deviations indicating that they are younger than the Pleiades. Therefore using an age range\footnote{10 Myr chosen as the low-end range based on the age of the youngest nearby moving group. 100 Myr chosen as the upper limit based on an extrapolation and comparison to Pleiades age objects.} of $<$ 10-100 Myr and spectral classifications of early-mid type L dwarfs, these objects have masses close to--or in some cases below-- the deuterium burning limit, making them exoplanet analogs. Since young brown dwarfs are nearby and isolated, they are ideal laboratories for detailed studies of cool, low-gravity, dusty atmospheres that are similar to directly imaged exoplanets. In this paper we examine the kinematic, photometric, and spectral features of the low surface gravity L5$\gamma$ dwarf 2MASSJ035523.37+113343.7 (2M0355). In section 2 we review published observations of 2M0355. In section 3 we describe new near-infrared spectral and imaging data, and in section 4 we evaluate indications of youth, including potential membership in nearby young moving groups. In section 5 we discuss the spectral energy distribution (SED) for 2M0355 as well as the near-infrared color-magnitude diagram for the brown dwarf population, highlighting the location of 2M0355 compared to directly imaged exoplanets. Conclusions are presented in section 6.
|
2M0355 is the reddest isolated L dwarfs yet characterized in the near- and mid-infrared. \citet{Cruz09} classified 2M0355 as L5$\gamma$, indicating low surface gravity spectral signatures. The similarity of the near-infrared spectrum to that of the $\sim$10 Myr planetary-mass object 2M1207b supports the conclusion that the object is young. Furthermore, a comparison with the near and mid-infrared colors of the known population of low surface gravity or L$\gamma$ dwarfs demonstrates that 2M0355 is the most extreme example of this class currently known. Combining optical spectra and absolute near to mid-IR photometry, we compared the full spectral energy distribution of 2M0355 with the field L5 dwarf 2M1507-16. We find that 2M0355 is underluminous in optical through $K$ band then switches to overluminous through at least 12$\mu$m compared to 2M1507-16. Calculating the bolometric luminosity by integrating over the optical and near-IR spectra as well as WISE photometry, shows that the overall luminosity of 2M0355 is overluminous compared to the field object. We conclude that enhanced photospheric dust, thought to be correlated with young, low-temperature, low-luminosity brown dwarfs and giant exoplanets, shifts flux to longer wavelengths creating the red SED. The position of 2M0355 on the near-IR color magnitude diagram supports this conclusion as it appears redward and underluminous of the known population in a similar region as 2M1207b and HR8799bcd. Combining new proper motion and parallax measurements we calculate UVW velocities to evaluate membership in nearby young moving groups. We find the kinematics consistent with the young thin disk and the UV velocities for 2M0355 place it in a busy part of velocity space for young objects. A careful kinematic comparison with nearby young groups and the field population leads us to conclude that 2M0355 has a 42\% chance of membership in AB Doradus. 2M0355 remains the brightest isolated low surface gravity L dwarf studied to date and will prove to be a useful comparative object in low-temperature atmosphere studies directly applicable to giant exoplanets. Despite the spectral similarity to 2M1207b in $H$ and $K$, 2M0355 is substantially different from the planetary-mass object in $J$band. This, combined with the older age estimate for 2M0355, cause the temperature and mass of 2M0355 to remain ambiguous. Nevertheless, we can use the object's absolute photometry and constrained age (assuming membership in AB~Doradus) to estimate these key properties. Using the evolutionary tracks for young, low mass objects of \citet{Baraffe02}, we estimate an effective temperature of $/sim$1500~K and a mass of $/sim$13 M$_{Jup}$ for an age of 50~Myr (the lower limit for the age of AB~Doradus). At the upper age limit for AB~Doradus, $/sim$150~Myr, the mass of 2M0355 would be closer to $/sim$30~M$_{Jup}$. As a field object, the absolute magnitudes of 2M0355 correspond to an object of $\sim$70~M$_{Jup}$, slightly below hydrogen burning minimum mass.
| 12
| 6
|
1206.5519
|
1206
|
1206.5802_arXiv.txt
|
{} {Grain growth has been suggested as one possible explanation for the diminished dust optical depths in the inner regions of protoplanetary ``transition" disks. In this work, we directly test this hypothesis in the context of current models of grain growth and transport.} {A set of dust evolution models with different disk shapes, masses, turbulence parameters, and drift efficiencies is combined with radiative transfer calculations in order to derive theoretical spectral energy distributions (SEDs) and images.} {We find that grain growth and transport effects can indeed produce dips in the infrared SED, as typically found in observations of transition disks. Our models achieve the necessary reduction of mass in small dust by producing larger grains, yet not large enough to be fragmenting efficiently. However, this population of large grains is still detectable at millimeter wavelengths. Even if perfect sticking is assumed and radial drift is neglected, a large population of dust grains is left behind because the time scales on which they are swept up by the larger grains are too long. This mechanism thus fails to reproduce the large emission cavities observed in recent millimeter-wave interferometric images of accreting transition disks.} {}
|
\label{sec:introduction} The evolution of circumstellar disks, the birthplaces of planets, is still enigmatic, even though the pioneering theoretical work on this topic was started almost 40 years ago by \citet{LyndenBell:1974p1945}. Viscous and/or gravitational stresses are the drivers of the disk accretion flows, which can be traced indirectly by the observed disk lifetimes and accretion rates \citep[e.g.,][]{Hartmann:1998p664,SiciliaAguilar:2006p10361,Hernandez:2007p4281,Fedele:2010p15973}. While the general trends of declining accretion rates and disk masses can be explained by viscous evolution, a sub-set of objects called transition disks remain mysterious. These objects appear dust-depleted in their inner regions, while the outer regions resemble normal circumstellar disks \citep{Strom:1989p9475, Skrutskie:1990p16132, Calvet:2002p10424, Espaillat:2007p17013, Espaillat:2010p17008, Andrews:2011p16142}. The sizes of the dust cavities range from a few to more than 70~AU \citep[e.g.,][]{Pietu:2006p17014, Hughes:2007p17475, Brown:2008p8893, Hughes:2009p17047, Andrews:2009p7729, Brown:2009p8895, Isella:2010p9438, Andrews:2010p17519, Isella:2010p17527, Andrews:2011p16142}. The gas content of the cavities is still largely unknown. It may also be reduced compared to the outer disk regions as found for example by \citet{Najita:2010p17082}, \citet{Dutrey:2008p17530}, or \citet{Lyo:2011p16794}; yet other works such as \citet{Pontoppidan:2008p9993} or \citet{Salyk:2011p17115} do detect gas inside the dust cavities. The time scale of this transition phase of disk evolution is estimated to be of the order of a few times $10^5$ years \citep{Skrutskie:1990p16132,Hartigan:1990p16133}. This estimate, however, is based on the transition disk frequency, which is still only a lower limit due to the lack of spatial resolution. Classifying transition disks only based on the SED can be misleading because steep decreases in the dust surface density can easily be missed due to the presence of small dust particles. Some mechanisms proposed to explain observations of transition disks include planet-disk interactions \citep[e.g.,][]{Rice:2003p15994,Zhu:2011p16181}, or photo-evaporation \citep{Clarke:2001p969,Alexander:2006p136,Ercolano:2008p13616}. Recent imaging of disks by \citet{Andrews:2011p16142} revealed a higher fraction ($>20$\%) of disks with large cavities for the mm-brightest sources. Higher fractions of transition disks, indicating longer disk clearing time scales, are more difficult to explain by photoevaporation. However, the dust emission signature of photoevaporating disks has not yet been self-consistently modeled, treating grain growth physics and dust-gas feedback. It is therefore unclear whether photoevaporating disks leave a dust rich or a completely dust and gas free cavity behind \citep[e.g.,][]{Alexander:2007p131,Garaud:2007p405}. The formation of large disk cavities by viscous evolution is generally problematic, even with the assumption that photoevaporation or some other mechanism is able to decouple the outer and inner disc at larger radii. Indeed the viscous time at 35~AU for typical values for the viscosity parameter\footnote{Assuming $\alphat = 3 \times 10^{-3}$. Both observational and theoretical works indicate values between $10^{-3}$ and $10^{-2}$.} is 3.4~Myr. On the other hand, even at 35~AU, the time scales for grain growth and radial drift are only a couple of thousand years. Any mechanism that triggers significant changes in the dust evolution (e.g., the emergence of a pressure maximum, see \citealp{Pinilla:2012p16999}) could therefore quickly induce observational signatures. The fact that grains grow and consequently become more mobile due to radial drift is likely an important part of the solution to this problem. Recent observations by \citet{Kraus:2012p16079} found a possible planetary-mass companion inside the cavity of the LkCa 15 transition disk. It remains to be shown whether planets cause gaps/pressure bumps or vice versa (e.g., \citealp{Kretke:2007p697,Brauer:2008p212}; Pinilla et al., in prep.). Determining the origin of transition disks is therefore one of the most fundamental issues in our efforts to forge a better understanding of planet formation. It has been suggested, although not demonstrated in any detail, that the growth and radial transport of dust could potentially explain the observed signatures of transition disks with large inner holes \citep[e.g.,][]{Dullemond:2005p378,Tanaka:2005p6703,Najita:2008p17130,Pontoppidan:2008p9993}. The observations indicate a decreased {\it optical depth} in the inner regions of transition disks both in the IR as well as in mm-observations, which does not necessarily imply diminished dust densities. Grain growth therefore offers two (related) pathways towards this end. First, particles decouple from the gas as they grow, which causes them to spiral inwards \citep{Weidenschilling:1977p865,Nakagawa:1986p2048}. This way, the dust optical depths are reduced by the actual removal of dust mass. Second, grain growth itself causes a decrease in the dust opacities, since larger particles emit less efficiently, which also decreses the otpical depth. Dust grains of sizes beyond a few centimeters become basically invisible to observations. However, the edges of transition disks seem to be relatively abrupt; they have been modeled as step functions \citep{Andrews:2011p16142,Isella:2012p17671} or steep power-law profiles \cite{Isella:2010p17527,Isella:2012p17671}. This suggests that the environment for grain evolution must abruptly change at this point in the disk. Pressure bumps or even gaps opened by planets might be possible explanations \citep{Lin:1986p12877,Zhu:2011p16181}. Another possibility could be the outer edges of dead-zones \citep{Gammie:1996p1515}. Dead zones are regions where the ionization fraction of the disk drops below the critical value needed to drive the magneto-rotational instability \citep[MRI,][]{Balbus:1991p4932}, the widely accepted source of turbulent viscosity. In this work, we want to investigate whether grain growth and transport -- either alone or aided by a dead zone -- can be the cause of the observed transition disk signatures. In Section~\ref{sec:background} we will discuss some of the equations which are crucial for the understanding of our modeling results. Section~\ref{sec:model} describes the model setups and assumptions. The resulting simulation outcomes and simulated observations are shown and explained in Section~\ref{sec:results}. Our findings are discussed and summarized in Section~\ref{sec:summary}.
|
\label{sec:summary} In this paper, we have investigated the ability of models of dust evolution to explain transition disks. We calculated simulated SEDs and 880~$\mu$m images from the output produced by a dust evolution code for a variety of different initial conditions. For clarity, we only considered effects directly induced by growth, fragmentation and transport of dust, \emph{without} pressure traps or planet induced gaps. The models were chosen to represent best-case scenarios for this mechanism to work, i.e. large gas disk masses and no viscous evolution were assumed. The only extensions to this model we have considered here are the ones which influence the grain size limits, i.e., the strength of turbulence and the gas surface density. We found that effects of grain growth can indeed produce dips in the SED such as found for many transition disks \citep[see also][]{Dullemond:2005p378}, however they fail to reproduce the cavities in the \mbox{(sub-)millimeter} images such as found by \citet{Brown:2009p8895,Andrews:2011p16142,Lyo:2011p16794}. Even in the extreme case where growth proceeds without any barriers and radial drift is switched off, no \textit{large} inner cavities can be formed within 5~Myrs of evolution. This leads to the conclusion that disks with large inner holes cannot be caused by grain growth \textit{alone}. Due to the fact that current models of disk photoevaporation \citep[e.g.][]{Owen:2010p14269} also fail to explain the inner cavities in the accreting objects, we propose that a combination of dust evolution with other effects such as pressure bumps, spiral arms or planet induced gaps could be the solution of the problem (see Pinilla et al., submitted to A\&A). Important questions to be answered from the observational side are: apart from the size of the cavity, is there is a distinct difference between disks with small and large cavities? What is the gas content inside the cavities? And do different sizes of dust show different cavity shapes in the same object \citep[see][]{Dong:2012p17966}? Theoretical models will need to investigate the trapping mechanisms which effectively shepherds the dust outside the cavity, irrespective of the particle sizes. Our findings can be summarized as follows: \begin{itemize} \item A grain size distribution for which the drift-induced grain size limit is smaller than the fragmentation induced size limit becomes inefficient in replenishing small dust due to a lack of fragmenting collisions. \item The time scale on which the remaining small dust is swept up by the largest, inward drifting grains is proportional to the distance to the central star and depends also on the dust-to-gas mass ratio and the turbulence strength. The dispersal of small dust is therefore from the inside out. \item Small amounts of small dust can be retained for several million years in the outer parts of the disk ($\gtrsim 10$~AU) if the dust-to-gas mass ratio is lower than a few times $10^{-4}$. \item Even if the inner regions are cleared of small dust via grain growth, most of the dust mass is still detectable by millimeter observations. Furthermore, the grain size is a smooth function of the stellocentric radius which means that the ``observable'' dust surface density decreases only slightly and the brightness profile stays approximately flat. This mechanism thus fails to reproduce the strong drop in millimeter emission which is typically observed in disks with large cavities. \item The model does not resemble the observations of large cavities, even if perfect sticking and no radial drift is assumed. The observed, sharp division between the inner and the outer regions call for a severe change in the disk properties. Pressure bumps that are strong enough to decouple both large and small dust from the accretion flow are most likely necessary to explain the observations. \end{itemize}
| 12
| 6
|
1206.5802
|
1206
|
1206.0750_arXiv.txt
|
We perform a study of 62 solar analog stars to compute their effective temperatures ($\teff$) using the Balmer line wing fitting procedure and compare them with $\teff$ values obtained using other commonly employed methods. We use observed H$\alpha$ spectral lines and a fine grid of theoretical LTE model spectra calculated with the best available atomic data and most recent quantum theory. Our spectroscopic data are of very high quality and have been carefully normalized to recover the proper shape of the H$\alpha$ line profile. We obtain $\teff$ values with internal errors of about 25\,K. Comparison of our results with those from other methods shows reasonably good agreement. Then, combining $\teff$ values obtained from four independent techniques, we are able to determine final $\teff$ values with errors of about 10\,K.
|
The effective temperature ($\teff$) is one of the most important parameters in the study of stars. For example, precise and accurate $\teff$ values allow us to reliably measure the chemical compositions of stars. Other important stellar parameters such as luminosity, radius, etc., can only be obtained once $\teff$ is known. A number of techniques have been devised to derive $\teff$. In this work, we use the relative flux level in the wings of H$\alpha$ line profile as an indicator of the star's effective temperature (e.g., Gehren 1981, Barklem et~al.\ 2002). In studies of stars like the Sun, systematic errors can be minimized if the data are carefully treated with a differential analysis. Thus, very precise $\teff$ values can in principle be derived using high quality data of solar analog stars. The aim of this work is to derive $\teff$ values using model fits to the H$\alpha$ line wings of 62 solar analogs and to compare the results with the $\teff$ values derived using three other methods.
|
Effective temperatures have been determined using the method of Balmer line fitting for a sample of 62 solar analog stars, with internal errors of about 25\,K. The other methods discussed in this work have internal errors of about 30 - 50\,K. The high precision of our $\teff$ values is useful to find small residual trends in the comparison with other methods. We find reasonably good agreement with the $\teff$'s obtained with the Ldr, IRFM, and R09 methods, but small trends and offsets for the residuals are detected and removed with linear corrections. We argue that high accuracy effective temperatures, with errors of order 10\,K, are possible to achieve for solar analog stars if several independent measurement are combined, mainly because the impact of errors is very small and can be understood and removed empirically.
| 12
| 6
|
1206.0750
|
1206
|
1206.6922_arXiv.txt
|
We present a set of three unified equations of states (EoSs) based on the nuclear energy-density functional (EDF) theory. These EoSs are based on generalized Skyrme forces fitted to essentially all experimental atomic mass data and constrained to reproduce various properties of infinite nuclear matter as obtained from many-body calculations using realistic two- and three-body interactions. The structure of cold isolated neutron stars is discussed in connection with some astrophysical observations.
|
\label{introduction} Neutron stars, born from the catastrophic gravitational collapse of massive stars with $M\gtrsim 8 M_\odot$ at the end point of their evolution, are among the most compact objects in the Universe. The extreme conditions encountered in their interior make their description a very challenging task. Indeed, neutron stars are expected to contain very different phases of matter, from ordinary nuclei to homogeneous nuclear matter; their core with densities exceeding several times the nuclear matter saturation density might contain additional particles like hyperons or even deconfined quarks~\citep{haensel2007}. The EDF theory allows for a consistent and computationally tractable treatment of these various phases. We have determined the global structure of neutron stars using three different \textit{unified} EoSs based on the recently developed EDFs BSk19, BSk20 and BSk21~\citep{goriely2010}.
|
\label{conclusion} We have developed a set of three different unified EoSs of cold isolated neutron stars in the framework of the EDF theory. These EDFs are all based on generalized Skyrme forces, fitted to essentially all experimental nuclear mass data with an rms deviation falling below 0.6 MeV~\citep{goriely2010}. At the same time, these EDFs were constrained to reproduce several properties of homogeneous nuclear matter (including the neutron-matter EoS) as obtained from microscopic calculations using realistic nucleon-nucleon potentials. These EoSs have been applied to compute the global structure of neutron stars. Even though our softest EoS BSk19 seems to be favored by measurements of the $\pi^-/\pi^+$ production ratio in heavy-ion collision experiments~\citep{xiao2009}, it is ruled out by the recently measured mass of PSR J1614$-$2230~\citep{demorest2010}. This conclusion might suggest that the dense core of neutron stars could be made of non-nucleonic matter~\citep{chamel2012}.
| 12
| 6
|
1206.6922
|
1206
|
1206.6113_arXiv.txt
|
Reaction cross sections of $^{169}$Tm($\alpha$,$\gamma$)$^{173}$Lu and $^{169}$Tm($\alpha$,n)$^{172}$Lu have been measured in the energy range $12.6\leq E_\alpha \leq 17.5$ MeV and $11.5\leq E_\alpha \leq 17.5$ MeV, respectively, using the recently introduced method of combining activation with X-ray counting. Improved shielding allowed to measure the ($\alpha$,$\gamma$) to lower energy than previously possible. The combination of ($\alpha$,$\gamma$) and ($\alpha$,n) data made it possible to study the energy dependence of the $\alpha$ width. While absolute value and energy dependence are perfectly reproduced by theory at the energies above 14 MeV, the observed change in energy dependence at energies below 14 MeV requires a modification of the predicted $\alpha$ width. Using an effective, energy-dependent, local optical $\alpha$+nucleus potential it is possible to reproduce the data but the astrophysical rate is still not well constrained at $\gamma$-process temperatures. The additional uncertainty stemming from a possible modification of the compound formation cross section is discussed. Including the remaining uncertainties, the recommended range of astrophysical reaction rate values at 2 GK is higher than the previously used values by factors of $2-37$.
|
\label{sec:mot} Two neutron capture processes, the $s$ and $r$ process, are required to produce the bulk of natural nuclides above Fe \cite{b2fh,kapgall,arngorr}. These two processes cannot, however, create 35 neutron-deficient, stable, rare isotopes between Se and Hg, which are termed $p$ nuclei. Photodisintegration of stable nuclei in the O/Ne shell of massive stars during a core-collapse supernova explosion has been suggested as a production mechanism for these nuclei \cite{arn,woohow,rayet95}. Such a so-called $\gamma$ process commences by sequences of ($\gamma$,n) reactions which are replaced by ($\gamma$,p) and ($\gamma$,$\alpha$) reactions when reaching sufficiently neutron-deficient nuclides in an isotopic chain \cite{raunic}. Two mass regions have remained problematic when explaining the production of $p$ nuclei by the $\gamma$ process in core-collapse supernovae: the lightest $p$ nuclei with mass numbers $A<100$ and those in an intermediate region at $150\leq A \leq165$ are underproduced \cite{arngorp,woohow,rayet95,rhhw02,hegXX}. While the explanation of the light $p$ nuclei most likely requires a different astrophysical model, the problem in the intermediate mass region may still be solved by improved reaction rates. For the $\gamma$ process, photodisintegrations happen in the plasma temperature range $2.0\leq T\leq 3.0$ GK. The temperature is tightly constrained by the necessity to photodisintegrate the lighter, more tightly bound seed nuclei while also retaining heavy $p$ nuclides. Several layers with slightly different temperatures contribute to $p$ nucleosynthesis in a star. The rates for heavier nuclei have to be known in the lower part of the temperature region because they would be destroyed completely at slightly higher temperatures. Moreover, ($\gamma$,$\alpha$) reactions have been found important for the intermediate and heavy mass region, whereas ($\gamma$,p) dominates in the lighter mass region of the $\gamma$-process \cite{raudeflect,rappgamma}. The photodisintegration reaction rates are usually computed from capture rates by applying the reciprocity principle of stellar rates \cite{fow74,raureview}. Therefore a measurement of $\alpha$ capture may determine also the photodisintegration rates, provided the g.s.\ contribution to the stellar rate is large. Due to the Coulomb barrier, the ($\alpha$,$\gamma$) reaction cross sections are tiny at astrophysical energy and thus currently unmeasurable. Going to as low energy as possible, however, already some discrepancies between data and predictions have been found in previous measurements. Low-energy $\alpha$ capture on heavy nuclei seems to be often overpredicted, although there is not yet enough data to draw a conclusive picture. It underlines, nevertheless, that the underproduction of p-nuclides in the range $150\leq A \leq165$ observed in stellar models may have a nuclear physics cause. Investigations of rates for the $\gamma$ process are not only important for nucleosynthesis in core-collapse supernovae. Simulations of the thermonuclear explosion of a White Dwarf (type Ia supernova) also found $p$ nuclei being produced in a $\gamma$ process \cite{travaglio}. Regardless of the site, $\gamma$ process studies require a sound determination of the relevant astrophysical reaction rates by nuclear physics investigations. The $\gamma$-process reaction networks include hundreds of nuclei and thousands of reactions, mostly on unstable nuclei. All of the reaction rates are predicted in the Hauser-Feshbach statistical model of nuclear reactions \cite{haufesh,adndt}. Since ($\gamma$,p) and ($\gamma$,$\alpha$) reactions occur at unstable isotopes, current experimental techniques have to aim for testing reaction model predictions at stability and to provide the data for a global improvement of these models and their input. The nucleus $^{169}$Tm is not a $p$ nuclide but it is close to the problematic mass range. Very few $\alpha$-induced reaction data are known in this mass range and none close to the astrophysically relevant energy range. This made $^{169}$Tm an interesting target for investigation using the newly introduced method of activation with subsequent X-ray counting, supplementing the convential $\gamma$-counting methods. Details of the experimental method and first results were already published in \cite{kis_plb,kis_npa}. Here we introduce the additional data, extending the ($\alpha$,$\gamma$) cross sections to lower energies, and focus on a discussion of the implications for constraining the astrophysical reaction rate for ($\alpha$,$\gamma$) at $\gamma$-process temperatures.
|
\label{sec:con} We have determined the reaction cross sections and astrophysical $S$ factors of $^{169}$Tm($\alpha$,n)$^{172}$Lu and $^{169}$Tm($\alpha$,$\gamma$)$^{173}$Lu at low energies, using our newly developed method of combining activation and X-ray counting. An improved shielding around the LEPS detector enabled us to measure the ($\alpha$,$\gamma$) reaction down to 12.6 MeV $\alpha$ energy, lower than before. The impact of the new results on the determination of the astrophysical reaction rates for $\alpha$ capture on $^{169}$Tm were discussed, using the full range of data. The combination of ($\alpha$,n) and ($\alpha$,$\gamma$) data was essential to disentangle errors in the predicted $\alpha$ width from those in other widths appearing in the reaction. It was found that the energy dependence of the $\alpha$ width below 14 MeV is different from what is expected by standard predictions. A modified, energy-dependent, local optical $\alpha$+nucleus potential was presented, able to describe both ($\alpha$,n) and ($\alpha$,$\gamma$) $S$ factors well across the measured energy range. Using the local potential, stellar ($\alpha$,$\gamma$) and ($\gamma$,$\alpha$) reactivities were calculated. They were found to be considerably lower at astrophysical $\gamma$-process temperatures than predictions using a standard optical potential. More speicifically, they were factors of $7.4-16.7$ below the rate calculated with the potential by \cite{mcf}. Ambiguities in the extrapolation to low energies, however, require an uncertainty of a factor of $2-3$ in the predicted rate, even when assuming that the shape of the energy dependence is understood. Further ($\alpha$,$\gamma$) measurements below the neutron threshold would be able to reduce this uncertainty. A further problem remains with identifying the nature of the potential modification. This leads to an even larger uncertainty in the calculated rate. Again, more low-energy data and an extended database including a wider range of nuclei would help to shed light on this issue. In conclusion, the recommended reactivity at a $\gamma$-process temperature of 2 GK is $0.06-0.95$ times the SMARAGD reactivity using the potential by \cite{mcf}, which translates to $2.3-37.0$ times the widely used standard values of \cite{adndt1}, thereby leading to an \textit{enhancement} in the ($\gamma$,$\alpha$) rate with respect to the values given in \cite{adndt,adndt1}. The case studied here is a good example for the restrictions and possible pitfalls that can be encountered when deriving astrophysical reaction rates from experimental data. It not only shows the importance of further measurements of reaction cross sections involving low-energy $\alpha$ particles to allow global studies of suitable optical potentials but also the significance of properly accounting for both, changing width sensitivities as well as thermally excited states, when interpreting the impact on an astrophysical reaction rate.
| 12
| 6
|
1206.6113
|
1206
|
1206.6263_arXiv.txt
|
In astrophysical systems, radiation-matter interactions are important in transferring energy and momentum between the radiation field and the surrounding material. This coupling often makes it necessary to consider the role of radiation when modelling the dynamics of astrophysical fluids. During the last few years, there have been rapid developments in the use of Monte Carlo methods for numerical radiative transfer simulations. Here, we present an approach to radiation hydrodynamics that is based on coupling Monte Carlo radiative transfer techniques with finite-volume hydrodynamical methods in an operator-split manner. In particular, we adopt an indivisible packet formalism to discretize the radiation field into an ensemble of Monte Carlo packets and employ volume-based estimators to reconstruct the radiation field characteristics. In this paper the numerical tools of this method are presented and their accuracy is verified in a series of test calculations. Finally, as a practical example, we use our approach to study the influence of the radiation-matter coupling on the homologous expansion phase and the bolometric light curve of Type Ia supernova explosions.
|
In studying astrophysical objects, a detailed understanding and description of the radiation field is vital, particularly if synthetic observables are to be computed for comparison with observations. Conceptually, the radiation field in a fluid is not independent of the fluid state and their co-evolution has to be described self-consistently within the framework of radiation hydrodynamics. Depending on the dynamical importance of the radiation field and the strength of the radiation-matter coupling, different strategies can be followed. If the energy associated with the radiation field is negligible compared to the total energy content, a de-coupled approach may be followed. For example, such a method has been used for the determination of synthetic light curves and spectra for Type Ia supernova (\snia) explosions around maximum light \citep[e.g.][]{Kasen2006, Kromer2009, Jack2011}. In cases where the radiative terms are dynamically important, however, a fully de-coupled treatment of the radiation field is not possible. For such applications a variety of different techniques have been used to follow the co-evolution of the radiation-matter state. In optically thick environments, the radiation field is well-described by the diffusion approximation and its evolution in radiation-hydrodynamical simulations can be incorporated by using flux limited diffusion methods \citep{Levermore1981}. This numerical prescription is used, for example, in the modelling of radiation dominated accretion discs \citep[e.g.][]{Turner2003, Hirose2009}. In the opposite case of low optical depth, the influence of the radiation field may be treated by including a radiative cooling term \citep[e.g.][]{Townsend2009, Marle2011}, as is often done in studies of stellar winds \citep[e.g.][]{Garcia-Segura1996, Mellema2002}. In the intermediate regimes between the two extremes, a full radiation-hydrodynamical description of the radiation-matter state is necessary, for example when accounting for convective motions in studies of stellar atmospheres \citep[e.g.][]{Stein1998, Asplund2000}, shock breakouts in supernovae \citep[e.g.][]{Blinnikov2000, Hoeflich2009, Piro2010} or when studying interactions of stellar explosions with circumstellar material \citep[e.g.][]{Fryer2010, kasen2010}. In this paper we present the numerical methods and the application of a new approach to radiation hydrodynamics that is based on Monte Carlo radiative transfer techniques. A similar strategy has been pursued in the calculations presented in \citet{Kasen2011}. Monte Carlo methods have already shown tremendous success in pure radiative transfer applications \citep[e.g.][]{Fleck1971, Abbott1985, Mazzali1993, Long2002, Carciofi2006, Kasen2006, Harries2011}. Within this probabilistic approach, complex radiation-matter interaction physics can be simulated and problems with arbitrary geometries can be addressed. Here we aim to extend the Monte Carlo method to radiation-hydrodynamical calculations and explore its practicality for modern astrophysical applications. The focus of this paper is to present the theoretical and numerical foundations and to verify the operation of our Monte Carlo radiation-hydrodynamical method. We begin with a brief overview of the theoretical concepts that govern radiating flows in Section \ref{sec:theory}, which is followed by an extensive description of the numerical methods of our approach in Section \ref{sec:numerics}. The physical accuracy and the computational feasibility of the techniques presented here are assessed in Section \ref{sec:testing}, in which the results of a series of test calculations are described. As a first application of the method in astrophysical environments we report in Section \ref{sec:application} on our investigation of \snias{} ejecta. In particular, we study the influence of the radiation-matter coupling on the ejecta structure and the resulting effects on the bolometric light curve during the homologous expansion phase. We summarize our results and conclude in Section \ref{sec:discussion}.
|
\label{sec:discussion} In this paper we presented a Monte Carlo approach to radiation hydrodynamical problems in astrophysical environments. By combining Monte Carlo radiative transfer methods that rely on the indivisible packet formalism \citep{Abbott1985} with the finite-volume hydrodynamical technique PPM \citep{Colella1984} we have aimed to retain the benefits of the Monte Carlo machinery for the modelling of complex interaction physics and arbitrary geometries. Here, our main focus lay on the development and presentation of the necessary numerical tools and on demonstrating the operation of this method, its physical accuracy and its computational feasibility. By using volume-based Monte Carlo estimators \citep{Lucy1999, Lucy2002, Lucy2003} in the reconstruction of radiation field characteristics, the maximum amount of information is extracted from the propagation behaviour of Monte Carlo packets and the Monte Carlo noise is minimized. A series of toy calculations has been performed to test the operation of the main components of our Monte Carlo radiation hydrodynamical method. In particular, the simulation of radiative shocks verified the accuracy of our approach to a standard radiation hydrodynamical test. As expected, due to the nature of the Monte Carlo method, calculations in optically thick environments are time-consuming but feasible and accurate, as the radiative shock examples showed. In general, all calculations were completed within hours to a day on a single desktop CPU. However, due to the very efficient scaling behaviour of the Monte Carlo algorithm to large numbers of processor cores, a future parallel implementation of the method provides the scope for significant decreases in the run time. The application to \snia{} ejecta successfully demonstrated the operation of our method to an astrophysical problem for which this method was primarily developed. In this exercise, the influence of the radiation-matter coupling on the density stratification during the near-homologous expansion phase of the ejecta has been investigated. The results we obtained are in agreement with the previous study by \citet{Woosley2007} who used the radiation hydrodynamics code \textsc{stella} \citep{Blinnikov2006}. The induced changes in the ejecta structure, however, were confirmed to have no significant influence on the bolometric light curve, as predicted by \citet{Pinto2000}. Despite the agreement of our results with previous studies, the \snia{} application also illustrated some difficulties of our approach. Our discretization scheme into packets of radiative energy allows us to easily construct all relevant radiation field characteristics and can be generalized to a fully frequency-dependent transport treatment in a straight-forward manner. However, in regions where the radiation field is close to LTE or to isotropy, the radiation force components are subject to considerable statistical fluctuations due to this discretization choice. Here, we suppressed this noise component by applying a smoothing kernel. Although beyond the scope of this work, in the future further reduction of the statistical noise should be explored by incorporating implicit Monte Carlo techniques \citep{Fleck1971} or a Monte Carlo radiative transfer approach that is based on the difference formulation \citep{Szoke2005, Brooks2005}. For such schemes, it will be important to consider how all the physical processes necessary to adequately address particular astrophysical applications can be implemented. As the main aim of this work was to establish the methods and the numerical framework of the Monte Carlo radiation hydrodynamics approach, all calculations were performed with a simplified implementation of the method. In particular, we restricted our tests to one-dimensional geometries and the radiative transfer was performed in a grey approximation with a very simple opacity prescription. However, we stress that these simplifications were made only to reduce the complexity and the computational effort of the test simulations. They do not affect the applicability and operation of the Monte Carlo radiation hydrodynamical approach itself. In the future, we aim to introduce more sophisticated opacity prescriptions and make the transition from grey transport to a fully frequency-dependent radiative transfer scheme. This will add a further level of sophistication to the method, but will not impact the operation of the radiation-matter coupling in our approach. The tools to realise the frequency-dependent transfer have already been developed and are provided for example in the framework of the macro-atom formalism by \citet{Lucy2002}. In addition to improving the physical accuracy of the radiative transfer, our long-term efforts will be directed towards a generalised implementation for multi-dimensional problems. With this generalization our radiation hydrodynamical method will include all major capabilities that have already made the Monte Carlo technique a very successful and rewarding approach for pure radiation transport applications.
| 12
| 6
|
1206.6263
|
1206
|
1206.2395_arXiv.txt
|
High redshift galaxies permit the study of the formation and evolution of X-ray binary populations on cosmological timescales, probing a wide range of metallicities and star-formation rates. In this paper, we present results from a large scale population synthesis study that models the X-ray binary populations from the first galaxies of the universe until today. We use as input to our modeling the Millennium II Cosmological Simulation and the updated semi-analytic galaxy catalog by \citet{Guo2011} to self-consistently account for the star formation history and metallicity evolution of the universe. Our modeling, which is constrained by the observed X-ray properties of local galaxies, gives predictions about the global scaling of emission from X-ray binary populations with properties such as star-formation rate and stellar mass, and the evolution of these relations with redshift. Our simulations show that the X-ray luminosity density (X-ray luminosity per unit volume) from X-ray binaries in our Universe today is dominated by low-mass X-ray binaries, and it is only at $z \gtrsim 2.5$ that high-mass X-ray binaries become dominant. We also find that there is a delay of $\sim 1.1\,\rm Gyr$ between the peak of X-ray emissivity from low-mass Xray binaries (at $z\sim 2.1$) and the peak of star-formation rate density (at $z\sim 3.1$). The peak of the X-ray luminosity from high-mass X-ray binaries (at $z\sim 3.9$), happens $\sim 0.8\,\rm Gyr$ before the peak of the star-formation rate density, which is due to the metallicity evolution of the Universe.
|
X-ray observations of galaxies are a key component to unraveling the physics of compact object formation. Stellar remnants and the distribution of the hot interstellar medium are best studied via their X-ray emission. With {\em Chandra} several global galaxy X-ray correlations have been established and appear to hold to at least $z=1$ \citep[e.g.][]{Bauer2002,Lehmer2008,Vattakunnel2012,Symeonidis2011,Lehmer2012}, implying a large-scale regularity to the production of X-ray binaries (XRBs) and hot gas, the primary sources of high energy emission in ``normal'' (non-AGN) galaxies. These include a very strong correlation between total X-ray emission from high-mass XRBs (HMXBs) and the galaxy-wide star-formation rate (SFR) \citep[e.g. ][]{Ranalli2003,Gilfanov2004a,Hornschemeier2005,Lehmer2010, Mineo2012a}, as well as a scaling between total X-ray emission from low-mass XRBs (LMXBs) and stellar mass ($M_*$) \citep{Gilfanov2004b,Lehmer2010,Boroson2011,ZGA2012}. These correlations present exciting challenges for models of accreting binary evolution \citep[e.g.][]{Fragos2008, Fragos2009, Belczynski2008} that predict significant variability in these relations based on, e.g., star formation history and metallicity. Recent ultradeep Chandra and multiwavelength surveys \citep[e.g. ][]{Xue2011} have permitted robust observational tests of these correlations in distant galaxies \citep[see also the review of][]{BH2005}. In a broad sense, curiously, the X-ray emission per unit SFR seems to evolve little with redshift, despite the evolution in gas metallicity, for example, over the same epochs \citep[e.g.][]{Lehmer2008}. \citet{Ptak2007} and \citet{TG2008} used rich multi-wavelength datasets to construct X-ray luminosity functions (XLFs) of normal galaxies in different redshift bins, and found that the XLF shows a statistically significant evolution with redshift and that it is the late type galaxies that are driving this evolution. Despite the significant investment of observing time in X-ray studies of distant normal galaxies, and the recent theoretical advances in modeling the X-ray source populations in nearby galaxies, our understanding of the cosmological evolution of populations of compact objects is still in its infancy. The first steps in this direction have been made by means of semi-analytic, empirical models \citep{WG1998,GW2001}. Considering a time-dependent SFR, these models predicted that the time required for binaries to reach the X-ray phase (due to the donor's nuclear evolution and angular momentum losses) leads to a significant time delay between a star-formation episode and the production of X-ray emission from X-ray binaries from this populations. State-of-the-art binary population models \citep[e.g.][]{Belczynski2008} provide us with a more physical picture of XRB populations as a function of galaxy properties. This type of detailed modeling has been applied to both starburst and elliptical nearby galaxies, where large populations of individual XRBs are resolved, and detailed ages, star-formation histories, and metallicities are measured. \citet{Belczynski2004} constructed the first synthetic XRB populations for direct comparison with the Chandra observed XLFs of NGC\,1569, a star forming dwarf irregular galaxy. \citet{Linden2009, Linden2010} developed models for the HMXB and Be XRBs of SMC, studying the XLF and the spatial distribution of this population, and investigating the effect of electron-capture supernovae of massive ONeMg stellar cores. \citet{Fragos2008, Fragos2009} performed extensive population synthesis (PS) simulations, modeling the two old elliptical galaxies NGC\,3379 and NGC\,4278, constraining the relative contribution to the observed XLF from sub-populations of LMXBs with different donor and accretor types, and the effects of the transient behavior of LMXBs. More recently, \citet{ZL2011} studied the evolution of XRB populations in late type galaxies, using a modified version of the BSE population synthesis code \citep{Hurley2002}. Considering different star-formation history scenarios, they were able to derive the X-ray emissivity as well as XLFs of the XRB populations. However, their model predictions are not in good agreement with observations for most of the Universe lifetime: from $4.5$ to $13.7\,\rm Gyr$ ($z=0-1.4$). The next step in this long-term effort is to apply these models in a cosmological context, in order to understand the nature of the accreting populations in high-redshift normal galaxies. In this paper, we study the global scaling relation of emission from XRB populations with properties such as SFR and stellar mass, and the evolution of these relations with redshift. More specifically, we developed a large grid of XRB PS models for which the information about the star-formation history and metallicity evolution were derived from cosmological simulations. Our models, which we have already constrained with observations of XRB populations of the local universe, give predictions for the evolution of the aforementioned global scaling relations back to the formation of the very first galaxies at redshift $z \gtrsim 15$. The plan of the paper is as follows. In Section~2 we describe our simulation tools, the {\tt StarTrack} PS code and Millennium II simulation, and the methodology we follow in developing models for the evolution of XRBs on cosmological timescales. Section~3 discusses the necessary bolometric corrections we employ in order to directly compare our model results with observations. In Section~4, we present the observational data we use to constrain our models and the statistical analysis we follow. We present the predictions of our maximum likelihood models for the evolution of XRB population in the high redshift universe in Section~5. Finally, in Section~6 we summarize the main findings and conclusions of our work.
|
We present here the first PS study of XRBs in a cosmological context. We created the largest grid of 288 PS models over nine metallicities, where we varied all the major model parameters that are known to affect the evolution of XRBs. We convolved these models with information about the star-formation history and metallicity evolution of the Universe from the Millennium II simulation and the \citet{Guo2011} semi-analytic galaxy catalog. The combination of our PS models with a cosmological simulation allowed us to derive global scaling relations of the XRB population with properties such as SFR and stellar mass, and the evolution of these relations with redshift. These predictions were compared statistically with observations from the local Universe in order to constrain our models and make robust predictions about the high-redshift Universe. The main conclusions of this work can be summarized as follows: \begin{enumerate} \item The statistical comparison of our model predictions revealed several models that are in excellent agreement with observations of the local Universe. Furthermore, we found that all models that are consistent with observational data in the local Universe exhibit very similar behavior at higher redshifts, indicating that our predictions for the redshift evolution of XRB populations are robust. \item Although the main goal of this paper is not to constrain the parameters of our PS models but instead to study the redshift evolution of XRB populations, our parameter study allows for some general conclusions on the possible values of the PS parameters we varied. The statistical analysis showed that a CE efficiency $\lambda\times\alpha_{CE}\lesssim 0.1$, where $\lambda$ is a measure of the central concentration of the donor star, and possibly a mixed initial mass ratio distribution (i.e. 50\% of binaries are drawn from a flat initial mass ratio distribution and 50\% from a ``'twin'' distributuon) is necessary in order to produce a LMXB population consistent with observations. These two parameters show some degeneracy, and a slightly lower CE efficiency would possibly favor a flat initial mass ratio distribution. On the other hand, model parameters that regulate the formation rate of BH-XRBs, such as the IMF and the stellar wind strength, appear to be important for both the LMXB and HMXB populations. This is not surprising as XRBs with BH accretors can be very luminous, and even a relatively small number of them can dominate the integrated X-ray luminosity of a population. Again, these parameters show a degeneracy as they affect the population in a similar way, and only their combined effect can be constrained from observations. \item The X-ray luminosity of XRBs per unit stellar mass or unit star-formation rate, is significantly higher for low-metallicity stellar populations, due to the fact that the young massive O/B stars, which are the precursors to compact object accretors in X-ray binaries, will lose less stellar mass from line-driven winds over their lifetimes for lower metallicities. This results in more numerous and more massive BHs, and subsequently more luminous X-ray binary populations. \item The X-ray emission from XRBs in the local Universe is dominated by LMXBs, and it is only at $z \gtrsim 2.5$ that HMXBs start to dominate the X-ray emission. The redshift that this transition happens is lower than that of the peak of star formation ($z\sim 3$), which shows that there is a delay between the formation of primordial binaries and the time at which these binaries become LMXBs. The formation timescale of LMXBs is estimated to be $\sim 1.1\,\rm Gyr$. This finding also implies that the X-ray luminosity of an XRB population in a ``typical'' galaxy, with a star-formation history similar to the global star-formation history of the Universe, is dominated by the X-ray emission of LMXBs. \item The formation rate of HMXBs follows closely the SFR. However, the X-ray luminosity from HMXBs per unit SFR ($L_{X,\, HMXBs}/SFR$) depends also on metallicity. Taking into account the metallicity evolution of the Universe, we find that the peak of the X-ray luminosity from HMXBs per unit volume, occurs $\sim 0.5\,\rm Gyr$ before the peak of the SFR density. Although the X-ray luminosity coming from HMXBs, per unit SFR, changes by half an order of magnitude over the evolution of the Universe, the observed signature of this effect is masked by the contamination of LMXBs, making the X-ray luminosity from the whole XRB population per unit SFR approximately constant with redshift. \item Finally, we found that at $z \lesssim 2.5$, where the emission from LMXBs dominates the total X-ray luminosity, the evolution of the X-ray luminosity per unit stellar mass is related to the average age of the LMXB population, while at $z \gtrsim 2.5$, the evolution of X-ray luminosity per unit stellar mass can be a probe of the evolution of the specific SFR with redshift. \end{enumerate} This first PS study of the evolution of XRB populations across comic time highlights the importance of the study of X-ray luminous normal galaxies at high redshift. Furthermore, it lays the ground for future work that will study role played by XRBs in the formation and evolution of galaxies through feedback processes \citep[e.g.][]{Mirabel2011,Zoltan2011,JS2012}.
| 12
| 6
|
1206.2395
|
1206
|
1206.5056_arXiv.txt
|
We revisit the causal backreaction paradigm, in which the need for Dark Energy is eliminated via the generation of an apparent cosmic acceleration from the causal flow of inhomogeneity information coming in towards each observer from distant structure-forming regions. A second-generation version of this formalism is developed, now incorporating the effects of ``recursive nonlinearities'': the process by which metric perturbations already established by some given time will subsequently act to slow down all future flows of inhomogeneity information. In this new formulation, the long-range effects of causal backreaction are damped, substantially weakening its impact for simulated models that were previously best-fit cosmologies. Despite this result, we find that causal backreaction can be recovered as a replacement for Dark Energy through the adoption of larger values for the dimensionless `strength' of the clustering evolution functions being modeled -- a change justified by the hierarchical nature of clustering and virialization in the universe, occurring as it does on multiple cosmic length scales simultaneously. With this and with the addition of one extra model parameter used to represent the slowdown of clustering due to astrophysical feedback processes, an alternative cosmic concordance can once again be achieved for a matter-only universe in which the apparent acceleration is generated entirely by causal backreaction effects. The only significant drawback is a new degeneracy which broadens our predicted range for the observed jerk parameter $j_{0}^{\mathrm{Obs}}$, thus removing what had appeared to be a clear signature for distinguishing causal backreaction from Cosmological Constant $\Lambda$CDM. Considering the long-term fate of the universe, we find that incorporating recursive nonlinearities appears to make the possibility of an `eternal' acceleration due to causal backreaction far less likely; though this conclusion does not take into account potential influences due to gravitational nonlinearities or the large-scale breakdown of cosmological isotropy, effects not easily modeled within this formalism.
|
INTRODUCTION: COSMIC CONCORDANCE AND CAUSAL BACKREACTION} One of the key questions in Cosmology today relates to the still-unsolved problem of what is causing the observed cosmic acceleration. Primarily indicated by Hubble curves constructed from luminosity distance measurements of Type Ia supernovae \citep{PerlAccel99,RiessAccel98}, this (possibly apparent) acceleration is just one aspect of the struggle for a consistent picture of the universe; a picture that would also require an explanation of the gap between the observed clustering matter content of $\Omega _\mathrm{M} \sim 0.3$ \citep[e.g.,][]{TurnerCaseOm0pt33} and the value of $\Omega _\mathrm{Tot} = 1$ as indicated by Cosmic Microwave Background measurements of spatial flatness \citep{WMAP7yrCosmInterp}, as well as a solution of the ``Age Problem/Crisis'' for matter-only (i.e., decelerating) cosmologies in which the universe appears to be younger than some of its oldest constituents \citep[e.g.,][]{TurnerCosmoSense}, along with explanations of other important issues. The standard approach to solving these problems is to introduce some form of ``Dark Energy'' -- the simplest case being the Cosmological Constant, $\Lambda$ -- which can fill the gap via $\Omega_\mathrm{DE} = \Omega_\mathrm{Tot} - \Omega_\mathrm{M} \sim 0.7$, which possesses negative pressure in order to achieve cosmic acceleration \citep[e.g.,][]{KolbTurner}, and which (for non-$\Lambda$ cases) recruits some form of internal nonadiabatic pressure \citep[e.g.,][]{CaldDDEsNotSmooth} in order to avoid clustering as matter does. Thus the introduction of Dark Energy (often using spatially-flat $\Lambda$CDM models) has led to a broadly-consistent ``Cosmic Concordance'' -- an empirical outline which has seemed so far to succeed fairly well \citep[e.g.,][]{WMAP7yrCosmInterp} at developing into a consistent cosmological picture. There are serious aesthetic problems with Dark Energy, however, as is well known; the most obvious being the problematical introduction of a completely unknown substance as the dominant component of the universe. Beyond that, a static (i.e., Cosmological Constant) form of Dark Energy suffers from two different fine-tuning problems: one being the issue that $\rho _{\Lambda}$ is $\sim$$120$ orders of magnitude smaller than what would be expected from the Planck scale \citep{KolbTurner}; and the other being a ``Coincidence Problem" \citep[e.g.,][]{ArkaniHamedCoinc}, questioning why observers {\it today} happen to live so near the onset of $\Lambda$-domination, given $\rho _{\Lambda}/ \rho _{\mathrm{M}} \propto a^{3}$. Moving to a dynamically-evolving Dark Energy (DDE), however, invites other problems, since the {\it self-attractive} nature of negative-pressure substances (i.e., $\partial E / \partial V = - P > 0$) means that a DDE may cluster spatially \citep{CaldDDEsNotSmooth}, possibly ruining it as a ``smoothly-distributed'' cosmic ingredient. This could potentially be solved through the ad-hoc addition of some form of nonadiabatic support pressure for the DDE \citep{HuGDM98}, but this is a possibility which we have argued against elsewhere \citep{BochnerAccelPaperI} on thermodynamically-based cosmological grounds. Besides Dark Energy, various other methods have been used to attempt to explain the observed acceleration, such as employing modified gravity \citep[e.g.,][]{TroddenModGravAccel}, or assuming the existence of an underdense void centered not too far from our cosmic location \citep[e.g.,][]{TomitaSNeVoid}. But to avoid the substantial (and perhaps needless) complications which arise when assuming departures from General Relativity, as well as the non-Copernican `specialness' implied by a local void, we will instead use the feedback from cosmological structure formation itself as a {\it natural} trigger for the onset of acceleration -- a trigger that automatically activates at just the right time for observers to see it, due to the fact that all such observers will have been created by that very same structure formation which generates the observed cosmic acceleration. This approach, known generally as ``backreaction'', was used by this author in \citet{BochnerAccelPaperI} (henceforth BBI) to find several clustering models which managed to precisely reproduce the apparent acceleration seen in Hubble curves of Type Ia supernova standard candles, while simultaneously driving a number of important cosmological parameters to within a close proximity of their Concordance values -- including the age of the universe, the matter density required for spatial flatness, the present-day deceleration parameter, and the angular scale of the Cosmic Microwave Background. The ability of our models to achieve these goals, despite the generally pessimistic view of backreaction typically held by researchers currently \citep[e.g.,][]{SchwarzBackReactNotYet}, was due to our adoption of an explicitly causal variety of backreaction, which admits the possibility of substantial backreaction from Newtonian-strength perturbations. The standard formalism used for computing backreaction effects, developed through the extensive work of Buchert and collaborators \citep[e.g.,][]{BuchertEhlers97,BuchertKerschSicka}, is non-causal in the sense that it drops all `gravitomagnetic' (i.e., velocity-dependent) effects, thus rendering it unable to account for metric perturbation information flowing (at the speed of null rays) from structures forming in one part of the universe, to observers in another. In similar fashion, typical studies of cosmic structure formation are also non-causal in that they use the Poisson equation without time derivatives of the perturbation potential, thus computing metric perturbations from {\it local} matter inhomogeneities only, disregarding all gravitational information coming in from elsewhere in space. The result is a mistaken (but widespread) notion that the entire Newtonian backreaction $Q_{N}$ can be expressed mathematically as a total divergence, thus ultimately rendering it negligible. But by restoring causality with a ``causal updating'' integral that incorporates perturbations to an observer's metric coming from inhomogeneities all the way out to the edge of their observational horizon, we find (BBI) that the sum of such `innumerable' Newtonian-strength perturbations -- which increase in number as $r^{2}$ within a spherical shell at distance $r$ from the observer, more than compensating for their $1/r$ weakening with distance -- adds up to a total backreaction effect that is not only non-negligible (regardless of the smallness of $v^{2}/c^{2}$ for most matter flows), but is in fact a dominant cosmological effect that is fully capable of reproducing the observed cosmic acceleration in a fully `concordant' manner. Despite these successes, a major problem with our model in BBI is its utter simplicity: it is clearly a toy model, with the results presented there serving primarily as `proof-of-principle' tests, rather than as precision cosmological predictions. Though the simplifications of the model are many, one in particular is serious in its consequences, while fortunately being not too difficult to fix: specifically, this is the dropping of what we have termed ``recursive nonlinearities''. Unrelated to {\it gravitational} nonlinearities, or to the nonlinear regime of density perturbations in structure formation, recursive nonlinearities embody the fact that the integrated propagation time of a null ray carrying perturbation information to an observer from a distant virializing structure would itself be affected by all of the other perturbation information that has already come in to cross that ray's path from everywhere else, during all times prior to arrival. In other words, causal updating is itself slowed by the metric perturbation information carried by causal updating, creating an operationally nonlinear problem. This issue was necessarily neglected in BBI, as that work was devoted to introducing our `zeroth-order' approach to causal backreaction. But here we fix this problem, incorporating recursive nonlinearities into a new, `first-order' version of our phenomenological model. We will find that this alteration significantly changes our results, causing a profound weakening of the backreaction effects generated by a given level of clustering, as well as significantly damping the long-term effects of information from old perturbations coming in from extreme distances. In order to retain causal backreaction as a viable model for generating the observed cosmic acceleration -- presuming here that this should indeed be done -- it will be necessary to re-interpret the meaning of our (inherently empirical) `clumping evolution functions' to now consider the effects of hierarchical clustering on a variety of cosmic scales. Doing this, we will show that a successful alternative concordance can once again be achieved, with the right amount (and temporal behavior) of acceleration, and with good cosmological parameters. This paper will be organized as follows: in Section~\ref{SecRecNonlinForm}, we will re-introduce our original causal backreaction formalism, and then describe the changes implemented in order to incorporate recursive nonlinearities into the model. In Section~\ref{SecRecNonlinResults}, we will explore the results of the new formalism, and discuss the implications of the model parameters that are now needed to achieve good data fits. Furthermore, we will discuss how the damping effects due to recursive nonlinearities would alter the key factors that determine the `ultimate' fate of the universe, as was discussed for our original formalism in BBI, given an acceleration driven by causal backreaction rather than by some form of Dark Energy. Finally, in Section~\ref{SecSummConclude}, we conclude with a summary of these ideas and results, highlighting the role of causal backreaction as a fundamental component of cosmological analysis and modeling.
|
SUMMARY AND CONCLUSIONS} In this paper, we have revisited the causal backreaction paradigm introduced in \citet{BochnerAccelPaperI}, for which the apparent cosmic acceleration is generated not by any form of Dark Energy, but by the causal flow of information coming in towards a typical cosmological observer from a multitude of Newtonian-strength perturbations, each one due to a locally clumped, virializing system. Self-stabilized by vorticity and/or velocity dispersion, such perturbations are capable of generating positive volume expansion despite their individually-Newtonian natures. Noting that previous `no-go' arguments against Newtonian-level backreaction are based upon non-causal backreaction frameworks, we see that the sum total of these small but innumerable perturbations adds up to an overall effect that is strong enough to explain the apparent acceleration as detected by Type Ia supernovae, as well as permitting the formulation of an alternative cosmic concordance for a matter-only, spatially-flat universe. Our purpose here has been to develop and test a second-generation version of this causal backreaction formalism, filling in one of the most important gaps of the original `toy model' by including what we have termed ``recursive nonlinearities'' -- specifically referring to the process by which old metric perturbation information tends to slow down the causal propagation of all future inhomogeneity information, therefore reducing the effective cosmological range of causal backreaction effects, and thus damping the strength of their overall impact upon the cosmic evolution and upon important cosmological observations. Utilizing the new simulation program introduced here, which now incorporates recursive nonlinearities into causal backreaction, we find profound differences in the resulting cosmological model calculations. For a given magnitude of self-stabilized clustering assumed for large-scale structure, denoted by dimensionless model input parameter $\Psi _{0}$, the overall power of causal backreaction is now considerably weaker, in addition to fading out relatively rapidly after the growth of clustering ceases. This is unlike the results of the original model, in which causal backreaction effects would continue to grow regardless of any late-time saturation of clustering, due to the causally-expanding ``inhomogeneity horizon'' seen by an observer which continually brings more `old' inhomogeneities into view from ever-greater cosmic distances. After discussion of some of the possible reasons for which causal backreaction may now appear to fall short of its cosmological goals -- either due to issues regarding the fundamental mechanism itself, or due to our highly simplified treatment of it -- we then considered a very straightforward way in which the paradigm may be fully recovered as a cosmological replacement for Dark Energy: all that is needed is the adoption of $\Psi _{0}$ values greater than unity. Though representing an ad-hoc modification of the original formalism, the change makes astrophysical sense in a number of ways. Rather than viewing the clumping evolution function $\Psi (t)$ as simply representing the fraction of cosmic matter in the `clumped' versus `unclumped' state at any given time, $\Psi (t)$ can now be recognized (more realistically) as representing the total backreaction effect of hierarchical structure formation in the universe, where clustering and virialization take place simultaneously on a number of different cosmic length scales -- from stellar clusters, to individual galaxies, to galaxy clusters, etc. Model input parameter $\Psi _{0} \equiv \Psi (t_{0})$ is now interpreted as the effective number of `levels' of completed clustering that exists (at current time $t_{0}$) in the large-scale structure when one sums over the (partial or total) clustering of matter on all relevant cosmic scales. Given this enlarged parameter space with $\Psi _{0} > 1$ now permitted, we once again find a selection of model cosmologies that succeed (despite the damping effects of recursive nonlinearities) at reproducing the observed cosmic acceleration, while also re-establishing an alternative cosmic concordance by producing output parameters that match the observables derived from several of the most important cosmological data sets. Furthermore, astrophysical considerations regarding the necessary {\it input} parameters for these apparently successful models -- specifically, the need to assume a sufficiently early beginning of clustering -- results in a preference by the new formalism for $\Psi (t)$ models that reflect the late, nonlinear phase of structure formation. This is an improvement over the old formalism without recursive nonlinearities -- which had preferred models that embody the early phase of clustering, with linearized matter fluctuations -- since it is this final, nonlinear stage of clustering during which virialization occurs via the generation of vorticity and velocity dispersion, and hence represents the more astrophysically reasonable source of substantial causal backreaction. Noting that the only problem still remaining for this new concordance is the somewhat excessively large amount of clustering required to achieve it -- that is, $\Psi _{0} \simeq 4$, rather than what we consider to be more reasonable values like $\Psi _{0} \sim 2 - 3$ -- we then determined that this problem could be successfully fixed (i.e., a good concordance generated with $\Psi _{0} < 3$) by introducing ``early saturation'', in which the clumping evolution function $\Psi (t)$ reaches its ultimate value of $\Psi _{0}$ somewhat in the past ($z \sim 0.25$), and then changes little thereafter. This is a highly reasonable adjustment to the formalism, since in the real universe ``gastrophysics'' feedback exists which creates superheated baryons, sending large amounts of material back into the intergalactic medium, thereby slowing down the continued clustering of matter at late times; not to mention the likely slowdown of clustering due to the feedback effects of the backreaction, itself. The only major drawback of this new feature is the addition of an extra model input parameter -- the epoch of saturation, $z_\mathrm{Sat}$ -- which results in a degeneracy within $(z_\mathrm{Sat},\Psi _{0})$-space, providing a range of models that all fit the Type Ia supernova data well, yet lead to significant differences for certain output cosmological parameters. The greatest variation in the output results due to this degeneracy occurs for the observable jerk parameter, $j_{0}^{\mathrm{Obs}}$, hence implying a loss of predictability for $j_{0}^{\mathrm{Obs}}$ by our causal backreaction formalism. This is a significant loss, given the previous findings from \citet{BochnerAccelPaperI} (without recursive nonlinearities) which had indicated that $j_{0}^{\mathrm{Obs}} >> 1$ was the most distinctive signature of causal backreaction, thus serving as the clearest way for distinguishing it from Cosmological Constant $\Lambda$CDM (or from anything close to it), since flat $\Lambda$CDM always requires $j_{0} = 1$. It thus becomes more difficult to find a falsifiable test of the causal backreaction paradigm, a test that is needed to definitively distinguish it from Dark Energy in order to eventually rule out one cosmological approach in favor of the other. Finally, concerning the `ultimate' fate of the universe, we note that the incorporation of recursive nonlinearities tends to shut down any strong apparent acceleration effects fairly quickly once the ongoing clustering (i.e., the continued growth of $\Psi (t)$) finally stops. Even more dramatic is the way in which the metric perturbation function, $I^{\mathrm{RNL}}(t)$, becomes essentially locked in place when approaching too close to unity, making it an even greater obstacle in terms of preventing the acceleration (apparent or otherwise) from completely taking over the cosmic evolution. This makes the scenario of a perpetual, `eternal' acceleration seem even less likely than it already did in \citet{BochnerAccelPaperI}; though the now-unbounded nature of $\Psi (t)$ could potentially provide some aid in producing a long-term acceleration, as long as virialized structure can continue to form on ever-larger cosmic scales, without any fundamental upper limit to the sizes of coherent structures. Furthermore, the question of the ultimate cosmic fate is once again complicated by the possible backreaction contributions of gravitationally nonlinear terms, and the (unavoidable) eventual breakdown of the approximation of the universe as ``smoothly-inhomogeneous'' -- both complications representing scenarios which our toy-model formalism is not presently designed to account for. In summary, we conclude that our causal backreaction formalism remains successful at generating an alternative cosmic concordance for a matter-only universe, without requiring any form of Dark Energy; though the necessary incorporation of recursive nonlinearities into these models implies that a significantly stronger amount of such backreaction than before is now needed, acting throughout the crucial `acceleration epoch' of $z \sim 0.2 - 2$ or so, in order to provide a degree of observed acceleration sufficient to match the cosmological standard candle observations.
| 12
| 6
|
1206.5056
|
1206
|
1206.0735_arXiv.txt
|
We present measurements of the specific ultraviolet luminosity density from a sample of 483 galaxies at $6 \lesssim z \lesssim 8$. These galaxies were selected from new deep near-infrared {\it Hubble Space Telescope} imaging from the Cosmic Assembly Near-infrared Deep Extragalactic Legacy Survey, Hubble UltraDeep Field 2009 and WFC3 Early Release Science programs. In contrast to the majority of previous analyses, which assume that the distribution of galaxy ultraviolet (UV) luminosities follows a Schechter distribution, and that the distribution continues to luminosities far below our observable limit, we investigate the contribution to reionization from galaxies which we can observe, free from these assumptions. Due to our larger survey volume, wider wavelength coverage, and updated assumptions about the clumping of gas in the intergalactic medium (IGM), we find that the observable population of galaxies can sustain a fully reionized IGM at $z =$ 6, if the average ionizing photon escape fraction (f$_{esc}$) is $\sim$30\%. A number of previous studies have measured UV luminosity densities at these redshifts that vary by a factor of 5, with many concluding that galaxies could not complete reionization by $z =$ 6 unless a large population of galaxies fainter than the detection limit were invoked, or extremely high values of f$_{esc}$ were present. The observed UV luminosity density from our observed galaxy samples at $z =$ 7 and 8 is not sufficient to maintain a fully reionized IGM unless f$_{esc}$ $>$ 50\%. We examine the contribution from galaxies in different luminosity ranges, and find that the sub-$L^{\ast}$ galaxies we detect are stronger contributors to the ionizing photon budget than the $L >$ $L^{\ast}$ population, unless f$_{esc}$ is luminosity dependent. Combining our observations with constraints on the emission rate of ionizing photons from Ly$\alpha$ forest observations at $z =$ 6, we find that we can constrain f$_{esc} <$ 34\% (2$\sigma$) if the observed galaxies are the only contributors to reionization, or $<$ 13\% (2$\sigma$) if the luminosity function extends to a limiting magnitude of M$_{UV} = -$13. These escape fractions are sufficient to complete reionization by $z =$ 6. Current constraints on the high-redshift galaxy population imply that the volume ionized fraction of the IGM, while consistent with unity at $z \leq$ 6, appears to drop at redshifts not much higher than 7, consistent with a number of complementary reionization probes. If faint galaxies dominated the ionizing photon budget at $z =$ 6--7, future extremely deep observations with the {\it James Webb Space Telescope} will probe deep enough to see them, providing an indirect constraint on the global ionizing photon escape fraction.
|
The reionization of the intergalactic medium (IGM) is the last major phase transition in the Universe, and as such has been a major topic of recent study. Over the past decade, a number of lines of evidence have yielded improved constraints on the duration of reionization, the completion redshift, and the primary source of ionizing photons. Observations of the Cosmic Microwave Background (CMB) measure the Thomson scattering optical depth due to electrons in the path. The updated 7 year {\it Wilkinson Microwave Anisotropy Probe} results measure $\tau =$ 0.088 $\pm$ 0.014, which corresponds to a redshift of instantaneous reionization of $z =$ 10.6 $\pm$ 1.2 \citep{komatsu11}. However, reionization is likely to be a much more extended event, especially if faint galaxies play a dominant role. Measurements of the kinetic Sunyaev-Zeldovich (kSZ) effect with the South Pole Telescope (SPT) from \citet{zahn11} have recently placed a limit on the duration of reionization, of $\Delta z <$ 7.9. When combined with the {\it Wilkinson Microwave Anisotropy Probe} ({\it WMAP}) constrains on instantaneous reionization, they conclude at 95\% confidence that reionization was complete at $z >$ 5.8. Additionally, observations of the \lya\ forest in high-redshift quasars also imply an end to reionization at $z \sim$ 6, as the measurements of the near-zones around $z =$ 6 quasars yield results consistent with a negligible neutral fraction \citep[e.g.,][]{fan06}, though more $z >$ 6 quasar sight-lines may be needed to conclusively verify this result \citep{mesinger10} \citep[see also][]{becker07}. Nonetheless, some sources must be responsible for reionizing the IGM between 6 $\lesssim$ z $\lesssim$ 14. Quasars were a natural choice, as they are extremely bright, and the bulk of their ionizing photons can escape. However, the quasar luminosity function peaks at $z \sim$ 2, and falls off rapidly toward higher redshift \citep[e.g.,][]{hopkins07}. Additionally, observations of the X-ray background rule out a dominant contribution to the reionizing photon budget from quasars \citep{dijkstra04b}. Population III stars, due to their predicted high masses \citep[e.g.,][]{bromm04,glover05}, were very efficient emitters of ionizing photons \citep{tumlinson00, bromm01, schaerer02}. Their overall contribution to reionization may, however, have been limited \citep[e.g.,][]{greif06}. This conclusion has been strengthened by recent simulation results which have corrected the mass scale of the first stars down to less extreme values \citep{clark11,greif11,greif12}. The most likely source for the bulk of ionizing photons is thus from star-forming galaxies themselves, where ionizing photons are created from the massive stars present during ongoing star formation. This has been well-studied observationally, but it was only in the last decade when large samples of galaxies close to the reionization epoch were compiled. Following wide and deep surveys with the Advanced Camera for Surveys (ACS) onboard the {\it Hubble Space Telescope} ({\it HST}), large samples of $z \sim$ 6 galaxies were compiled \citep[e.g.,][]{giavalisco04,bunker04}. While \citet{giavalisco04} found that the specific UV luminosity density (and thus the star-formation rate density) was roughly constant from $z =$ 4 to 6, \citet{bunker04} found, via deeper observations, that this quantity appeared to decline towards higher redshift. Using the commonly accepted (at the time) large value for the clumping factor of the IGM, they concluded that galaxies were unable to account for the necessary ionizing photons to reionize the universe. \citet{yan04} pointed out that if the faint-end slope of the galaxy luminosity function was steep, it could be that $L < 0.1L^{\ast}$ galaxies dominated the ionizing photon budget, and thus allowed galaxies to complete reionization by $z =$ 6. Regardless of the interpretation, these galaxy samples contained only $\sim$ 50-100 galaxies, each of which was only detected in one band, given the lack of deep {\it HST} near-infrared observations at the time. The advent of the Wide Field Camera 3 (WFC3) on {\it HST} has opened the door to the $z \geq$ 6 universe, allowing the first investigation into whether reionization completed at $z =$ 7 or earlier. This is combined with recent results which have vastly changed the expected IGM clumping factor to much lower values. \citet{finkelstein10a} investigated the combined UV luminosities of their observed $z =$ 7 galaxies, and found that while they could come close to reionizing the IGM, either fainter galaxies or high ($\geq$ 50\%) escape fractions were necessary to complete reionization by $z =$ 7 \citep[see also][]{bunker10} A number of studies have used measured luminosity functions of $z \sim$ 7 and 8 galaxies to infer their contribution to reionization, by integrating them down to an assumed magnitude limit. The majority of these studies found that, once galaxies fainter than the detection threshold are accounted for, galaxies could reionize the universe at such early times \citep[e.g.,][]{oesch10,mclure10,lorenzoni11,bouwens11d} \citep[see also][]{grazian11}. This is however reliant on a number of assumptions. First, the luminosity function results are susceptible to the assumption of a Schechter function parameterization. At $z \sim$ 5--6, where the samples are larger, there is no strong evidence for deviation from this function \citep[e.g.,][]{bouwens07,mclure09}. However, as we push closer to the Big Bang, galaxies are changing rapidly, thus at some point we may encounter an epoch where the Schechter function is no longer an accurate representation. This could be due to a variety of effects, with one being a lack of active galactic nuclei (AGN) feedback if AGNs are not yet present in the centers of all galaxies at very high redshifts, which is plausible depending on the speed with which supermassive black holes form in the early universe. Additionally, while recent evidence has indicated a very steep faint end slope at $z \geq$ 7 \citep[e.g.,][]{bouwens11, oesch12, bradley12}, the uncertainty on these measurements are large \citep[e.g., $\sigma (\alpha) \approx$ 0.2 at $z =$ 8][]{bradley12}. Finally, when integrating the luminosity function, one needs to choose a limiting magnitude, as suppression from the UV background will result in the gas in galaxies becoming heated below some circular velocity (i.e. mass) limit. Galaxies above this limit are dense enough where collisional excitation of H\,{\sc i} by electrons can overcome photoionization heating, and have their gas cool and form stars. However, we have no observational evidence for this value, and theoretical results from the literature yield values of $-$15 $< M_{lim} < -$10 \citep[e.g.,][]{finlator11b,munoz11,kulkarni11,choudhury08}. With a steep faint-end slope, a difference of this level can result in a difference in the integrated luminosity density by more than a factor of two. Here we measure the specific luminosity density of the observable galaxy population using the largest sample of $6 < z < 8$ galaxies yet compiled, which allows us to study the contribution of galaxies to reionization without the uncertainties inherent in invoking a parametrized luminosity function. This allows us to assess whether a significant contribution from galaxies below the observational limits is necessary to complete reionization at a given redshift. We emphasize our results at $z =$ 6, where we have a large sample, and each galaxy is detected in four individual imaging bands, yielding much more robust results over the previous studies which had access only to optical ACS data. In \S 2 we describe the datasets used in this study, as well as our photometry and sample selection methods. We discuss how we measured the rest-frame UV specific luminosity density, and corrected it for incompleteness down to our magnitude limit in \S 3, while in \S 4, we discuss the implications of our results on the reionization of the IGM. Throughout this paper we assume a concordance cosmology, with H$_\mathrm{o}$ = 70 km s$^{-1}$ Mpc$^{-1}$, $\Omega_m$ = 0.3 and $\Omega_\lambda$ = 0.7. All magnitudes are reported in the AB system, where $m_\mathrm{AB} = 31.4 - 2.5\log(f_\nu / 1\, \mathrm{nJy})$ \citep{oke83}. In the remainder of the paper we make use of luminosity functions from the literature over the redshift range of our study. We use the luminosity functions from \citet{bouwens07} for $z =$ 4, 5 and 6; \citet{bouwens11} for $z =$ 7; \citet{bradley12} for $z =$ 8 and \citet{bouwens11c} for $z =$ 10. These represent the most up-to-date luminosity functions for these redshifts at the time of this writing. The characteristic magnitude values from these studies are: $M^{\ast}$ = $-$20.98 ($z =$ 4), $-$20.64 ($z =$ 5), $-$20.24 ($z =$ 6), $-$20.14 ($z =$ 7), $-$20.26 ($z =$ 8) and $-$18.3 ($z =$ 10). \vspace{5mm}
|
\begin{itemize} \item{We measure $\rho_{UV}$ at $z =$ 6 to be 0.79 $\pm$ 0.06 $\times$ 10$^{26}$ erg s$^{-1}$ Hz$^{-1}$ Mpc$^{-3}$ for the observed population of galaxies; this is much more robust than previous measures, which had much smaller samples and were limited to single band detections, and differed by a factor of $>$ 5. Additionally, this result is independent of a Schechter function parametrization for the form of the luminosity function. Assuming C/f$_{esc} \sim$ 10, we find that the observed galaxy population at $z =$ 6 is sufficient to sustain a fully reionized IGM. In contrast to many previous studies, this result does not need to invoke galaxies fainter than our detection threshold.} \item{We measure $\rho_{UV}$ of the observed galaxy population at $z =$ 7 and 8 to be 0.44 $\pm$ 0.04 and 0.27 $\pm$ 0.06 $\times$ 10$^{26}$ erg s$^{-1}$ Hz$^{-1}$ Mpc$^{-3}$, respectively. These values fall short of what is needed to reionize the IGM. While fainter galaxies may render the population capable of completing reionization, even {\it JWST} will not go deep enough to see the necessary limiting magnitude at $z =$ 8.} \item{If we only consider the observed galaxies, we find a scenario where though $z =$ 7 and 8 galaxies contribute to reionization, the IGM only becomes fully reionized by $z =$ 6. This is consistent with independent studies of both \lya\ emission from normal galaxies, and the study of \lya\ absorption in the near-zone of the $z =$ 7 quasar, both of which find that the IGM at $z =$ 7 likely has a neutral fraction $\geq$ 10\%.} \item{Accounting for constraints on the total emission rate of ionizing photons from measurements of the \lya\ forest at $z =$ 6, we find that if our observed galaxies represent the total galaxy population, then the average $z =$ 6 escape fraction must be f$_{esc}$ $<$ 34\%. However, fainter galaxies likely do exist, and if the luminosity function extends down to M$_{lim} = -$13, then the average escape fraction constraint tightens to f$_{esc}$ $<$ 13\%. These escape fractions can sustain reionization $z =$ 6 either with our observed galaxies, or with the integrated luminosity function, for values of the clumping factor $\leq$ 3.} \item{We investigate the evolution of the volume ionized fraction x$_{HII}$ of the IGM. We do this by combining the observed galaxy population of our study and the luminosity functions from the literature with limits on the escape fraction considering both limiting magnitudes. While the IGM appears to be fully ionized by $z \leq$ 6, the volume ionized fraction appears to drop below unity by $z =$ 7, consistent with a number of complementary analyses. Considering only the (roughly known) contribution from galaxies at $z \geq$ 8, the volume ionized fraction may drop substantially below unity, though this is in mild tension with reionization models which incorporate the optical depth due to electron scattering. Better constraints on the $z \geq$ 8 luminosity functions, likely from {\it JWST}, will place stronger constraints on the reionization history at that early time. If the strong drop in the luminosity density at $z >$ 8 is verified, it may imply some additional source of ionizing photons at $z >$ 10 is warranted to remain consistent with the {\it WMAP}-measured Thomson scattering optical depth.} \end{itemize} We conclude that in order to make progress on the issue of when and how reionization happened, we need better constraints on the escape fraction of ionizing photons or the limiting magnitude of the UV luminosity function. As these quantities are correlated due to the constraints from the \lya\ forest, constraints on one will yield more robust constraints on the other. While direct measurement of the escape fraction at high redshift is not possible, much deeper observations with {\it JWST} should confirm whether the luminosity function extends down to at least M$_{lim} = -$15.5, which will place stronger constraints on the escape fraction at $z =$ 6, and uncover whether the $z =$ 7 galaxy population is sufficient to reionize the universe.
| 12
| 6
|
1206.0735
|
1206
|
1206.0818_arXiv.txt
|
Laser frequency noise is a dominant noise background for the detection of gravitational waves using long-baseline optical interferometry. Amelioration of this noise requires near simultaneous strain measurements on more than one interferometer baseline, necessitating, for example, more than two satellites for a space-based detector, or two interferometer arms for a ground-based detector. We describe a new detection strategy based on recent advances in optical atomic clocks and atom interferometry which can operate at long-baselines and which is immune to laser frequency noise. Laser frequency noise is suppressed because the signal arises strictly from the light propagation time between two ensembles of atoms. This new class of sensor allows sensitive gravitational wave detection with only a single baseline. This approach also has practical applications in, for example, the development of ultra-sensitive gravimeters and gravity gradiometers.
| 12
| 6
|
1206.0818
|
Subsets and Splits
No community queries yet
The top public SQL queries from the community will appear here once available.