adsabs/WIESP2022-NER huggingface dataset:
Paper: 2019ApJ...871...47P (fulltext)
the presence of perturbations is at (for details about that, see Pereira et al. 2018 ). From Equations ( 7 ) and ( 12 ), one has thatwhere use has been made ofas well as Equation ( 4 ) and . The perturbed surface tension , with the help of Equations ( 8 ) and ( 11 ), can be written asTherefore, likewise to the slow phase conversion case, no thin shell surface degrees of freedom are induced in perfect fluids when the phase conversion is rapid. Now we investigate solid hybrid stars. In this case, as commented in the previous section, . Besides, now changes with respect to its perfect fluid counterpart, Equation ( 17 ), and it is given by (a mere replacement of by )where summarizes the contribution to the radial shear stresses in the presence of perturbations (see Penner et al. 2011 and Krüger et al. 2015 for realizations of them). For slow phase conversions in solid hybrid stars, due to Equations ( 11 ) and ( 13 ), it still follows that and . However, nontrivial results hold for solid hybrid stars under rapid phase conversions. Indeed, for , we haveBefore calculating the thin shell surface tension for this case, let us expand Equation ( 13 ), because . From our assumption to elastic aspects in hybrid stars, one can generically write . Thus, after some simple calculations it follows from Equations ( 4 ), ( 13 ), and ( 23 ) that (recall that )which turns out to be exactly the same as that for slow reactions in solid stars (because there and ). From the above equation, one can clearly see that indeed in general. (Equation ( 24 ) also makes it clear that the Lagrangian change of the pressure is null for perfect fluids, as already known from other methods; Pereira Rueda 2015 ; Pereira et al. 2018 .) Now, for the thin shell surface tension induced by rapid reactions in solid hybrid stars, from Equations ( 12 ), ( 22 ), and ( 24 ), one obtains that The fact that rapid conversions in solid stars induce thin shell surface degrees of freedom could have interesting consequences; for instance, the induction of a surface mass (general relativistic energy) on a phase-splitting surface. Since the induced thin shell surface energy density is a local quantity, so is the shell’s proper mass (note that for first-order quantities, such as those associated with surface degrees of freedom, it is irrelevant to use R or R 0 for functions multiplying them)Total mass measurements at infinity will have to take into account the correction factor (related to the gravitational redshift), as we shall show in Equation ( 31 ). From Equations ( 7 ), ( 26 ), and ( 5 ), it follows thatGiven that , when only linear terms in ξ are kept, the above equation can be simplified to . Thus, from Equations ( 26 ) and ( 27 ),where it has been assumed that , m 0 being the background mass function ( because the background has been assumed not to have surface degrees of freedom—see the Appendix ) and m 1 being its first-order correction in the presence of perturbations. In summary, among other
Paper: 2020MNRAS.491.5826K (fulltext)
fourth equation of 11 ), to derive j x , 1, out . Then, if we take into account ( B6 ) and the third equation of ( B4 ), we get: where To obtain the first amendment N 1, out , we apply the fact that . Then, we take account ( B9 ), and N 1, out takes the following form: where and . N 1, ihs can be derived if we use the flow definition for the HS medium j x , 1, ihs that is proportional to and take into account ( B6 ). Then, we obtain: The first amendment N 1, ihs can be presented as follows: where . Here, the continuity condition of densities at HP, i.e. the condition N 0, ihs ( x hp , η) + N 1, ihs ( x hp , η) = N 0, out ( x hp , η) + N 1, out ( x hp , η) was applied (the second equation of 11 ). After substituting the proposed form in ( B12 ), we obtain: Due to the fact that the solution of the homogeneous equation ( B14 ) has the form and taking into account the condition where Φ( x = x hp , η) = 0, it is necessary to search for an Φ( x , η) in the following form: Consequently, by taking into account ( B15 ) and ( B14 ), Φ( x , η) takes the following form: where Note that S 2 ( x ), S 3 ( x ), and S 4 ( x ) were obtained in consequence of expanding of the exponent from ( B15 ) in a Taylor series and integrating appropriate expressions to derive Φ( x , η). Consequently, the first amendment N 1, ihs ( x , η) takes the following form: Now, we use, on the one hand, the flow definition for the inner region of the heliosphere j x , 1, in what is proportional to and, on the other hand, the value of its expression that is determined by ( B5 ). Then, we obtain Let us present the first amendment N 1, in in the following form: where , i.e. the continuity condition of densities at TS, namely N 0, in (1, η) + N 1, in (1, η) = N 0, ihs (1, η) + N 1, ihs (1, η) is applied (the first equation of 11 ). After substituting ( B19 ) in ( B18 ), we obtain The solution of the homogeneous equation ( B20 ) has the form , this means that a solution for Υ( x , η) should be sought in the following form: As a result, by taking into account the form ( B21 ) and the expression ( B20 ), we get Υ( x , η). After substituting it in ( B19 ), we finally obtain the first amendment N 1, in : where . APPENDIX C: SOLUTION FOR <inline-formula><tex-math id="TM0132" notation="LaTeX"></tex-math></inline-formula> BY AI To obtain a solution for the problem, we introduce the following dimensionless variables: , , , , and . Note that for such a problem, the diffusion coefficient κ can be rewritten in the following form: . This follows from the fact that and . Taking into account the system of ( 9 ) and restrict by a term N 0 in the first two equations of ( 11 ), we obtain a system of equations for finding the zero-order solution for each region of CR modulation: The last system can be rewritten in the following form: Then, the solution of the last system has the form To obtain the first amendments N 1, in , N 1, ihs , N 1, out , we insert ( C3 ) into ( 10 ): To derive j x , 1, in and j x ,
Paper: 2020ApJ...889..132G (fulltext)
lines to fade from TPSB to QPSB. Figure 3. Stacked spectra for SF (top), TPSB (middle), and QPSB (bottom). All the spectra are normalized at 4100 Å. The shadow marks the region of H δ . 2.4. SFR, Gas-phase Metallicity, and Σ<sub>H <sc>i</sc></sub> The emission lines may be contaminated by active galactic nuclei (AGNs) and shocks, which would lead to the overestimation of the star formation rate (SFR) when we use H α as tracer of star formation. To this end, we investigate the excitation mechanisms for SF and TPSB regions according to the traditional BPT diagram (Baldwin et al. 1981 , left panel of Figure 4 ). In this diagram, we impose the S/N of these four emission lines (i.e., H α , [N ii ], H β, and [O iii ]) to be greater than 3. For the spaxels with H α and [N ii ] S/N larger than 3, we use the relation between the EW (H α ) and [N ii ]/H α (right panel of Figure 4 ) as the diagnostic diagram of excitation mechanisms (Cid Fernandes et al. 2010 ). We find no spaxels in the Seyfert region, which indicates that PGC 26218 does not host an AGN. Although some spaxels locate at the composite and LINER regions, the effect from shock is ignorable considering that most of the spaxels locate at the SF region and the contribution from shocks to emission lines could be low in the SF region. Taking into account that all of the SF and TPSB spaxels have EW(H α ) −3 Å and the percent light contribution (at λ = 5100 Å) of young populations in those spaxels is larger than 4%, we believe that most of the contribution is dominated by ionization due to young stars and only a tiny fraction by hot low-mass evolved stars/post AGB stars (González Delgado et al. 2016 ; Lacerda et al. 2018 ). At any rate, our SFR value is an upper limit. Figure 4. Left panel: the traditional BPT diagnostic diagram for SF and TPSB regions. The blue squares and red triangles represent the spaxels with S/N of four emission lines larger than 3 for SF and TPSB, respectively. The dashed and solid lines are from Kauffmann et al. ( 2003a ) and Kewley et al. ( 2001 ), respectively. Right panel: an alternative diagnostic diagram of excitation mechanisms for SF and TPSB regions. The hollow squares and triangles represent the spaxels with S/N of H β or [O iii ] lower than 3 for SF and TPSB, respectively. The blue solid squares and red triangles are the same as in the left panel. The horizontal dashed line is from Kewley et al. ( 2006 ) and the vertical dashed line is from Kewley et al. ( 2001 ). In order to estimate the spatially resolved SFR from extinction-corrected H α emission lines, we calculate the extinction according the Balmer decrement assuming a Cardelli et al. ( 1989 ) extinction curve with case B condition for each Voronoi bin. The extinction-corrected SFR is calculated using the formula given by Kennicutt ( 1998a ):where L (H α ) is the extinction-corrected H α luminosity. For the total extinction-corrected SFR, we construct a spectrum by stacking all spaxels in Voronoi bins
Paper: 2020MNRAS.499.2245C (fulltext)
ratios compared to the theoretical values expected from Doppler boost model. In three of the four candidates (J0247, J0249, and J0254), the observed values show deviations from the expected values at the ≳2σ level in at least two band pairs with p-values of 0.189, 0.013, and 0.113. In J0246, the observed values are generally consistent with those expected from Doppler boost. Figure 4. Observed frequency-dependent variability amplitude ratios compared with the expected values from the Doppler boost model. Different colours represent different combinations of band pairs. The black lines represent the one-to-one relation. The error bars denote 1σ uncertainties. See Section 5.2 for details. Fig. 5 shows the allowed parameter space for the Doppler boost model to explain the observed multiband variability amplitudes (Table 4 ). The parameters considered are the total black hole mass M BH , binary orbital inclination angle i , the fraction of the total emission from the secondary black hole f 2 , and the mass ratio q . Our other fiducial parameters are the rest frame orbital period P orb = P GLS /(1 + z ) and α ν in griz bands (Table 4 ). Figure 5. Doppler boost model parameter space estimates for the four candidate periodic quasars. The four sub-panels for each quasar represent griz bands. The dashed (shaded) contours denote f 2 = 1.0 ( f 2 = 0.8), where f 2 is defined as the fraction of the total emission from the secondary black hole. Different colours represents different mass ratios with q = 0.0, 0.05, 0.1, and 0.2 for blue, green, orange, and red, respectively. The black vertical line with the grey-shaded region shows the virial mass estimated from SDSS spectra (Fig. 3 ) and its 1σ statistical error. The orbital inclination angle i = 90° for an edge-on view and i = 0° for a face-on view. For J0246, while the observed frequency-dependent variability amplitudes are generally consistent with the expected values under the Doppler boost hypothesis (Fig. 4 ), there is almost no allowed parameter space unless the following conditions are all met: 1. The virial black mass is significantly underestimated, even after considering a 0.4-dex systematic error (e.g. Shen 2013 ), 2. A high fraction (e.g. ) of the total emission is from the secondary black hole fueled by its mini accretion disc, and 3. The binary system is close to being an edge-on view. For the other three quasars, the observed multiband variability amplitudes are generally consistent with the expected values under the Doppler boost hypothesis, although there is a stringent requirement on the allowed parameter space. Liao et al. ( 2020 ) have tested the Doppler boost hypothesis for J0252. They suggest that the Doppler boost model is disfavoured for J0252 based on both the frequency-dependent variability amplitude ratios and the multiband variability amplitudes. In summary, there is evidence against the Doppler boost hypothesis in all five periodic candidates in our parent sample based on the
Paper: 2017MNRAS.468..378S (fulltext)
The results we obtain span 7 10 7 2 10 9 M . 3.3 X-ray luminosities and Eddington fractions We returned to the Swift XRT X-ray spectra we presented in S11 in light of the new redshift measurements, and fitted each spectrum with an absorbed power-law model in xspec , making use of the most up-to-date Galactic column densities (Willingale etal. 2013 ) and adopting Cash statistics. Swift XRT observations additional to our fill-in programme were available for one source, XMMSL1J002202.9254004. We included these in a combined X-ray spectrum totalling 12ks of exposure, created using the method of Evans etal. ( 2009 ) and after establishing no significant changes in spectral hardness had occurred. From the XRT best fits, we obtained the observed 210keV X-ray luminosities of our AGN. We estimated the Eddington luminosities, L Edd , using our mass estimates in the equation ergs 1 . To obtain the Eddington ratio, the bolometric luminosity is required. Our measured X-ray luminosities significantly underestimate the total output, by an approximate factor of 10 (Elvis etal. 1994 ). Marconi etal. ( 2004 ) derived a bolometric correction to convert from 210keV X-ray luminosity to L Bol (their equation 21), which we adopt here. The resulting bolometric luminosities and Eddington ratios are listed in Table 4 . Table 4. Estimates of black hole mass and Eddington ratio. Optical photometry from USNOB1.0 ( R , I ) or NOMAD ( V ) is given for the observed band closest to rest frame 5100 . Black hole mass estimates use the virial method described in Shen etal. ( 2011 ) and Vestergaard Peterson ( 2006 ) (H linewidth taken from SDSS for this source). Bolometric luminosity estimates, L Bol , use the corrections given in equation 21 of Marconi etal. ( 2004 ). Most of these AGN have X-ray luminosities that imply a bolometric luminosity around a tenth of a per cent Eddington (mean value 0.2). The spread is large, however; our least luminous AGN, XMMSL1J131651.2084915 with L 210keV 510 42 ergs 1 , is emitting at a few hundredths of the Eddington rate, whilst our most distant source, XMMSL1J095336.4161231 at z 0.873, may be emitting at or close to the Eddington rate. 3.4 -ray and radio searches At the hard X-ray/-ray energies of 15150keV, none of our sources are detected in the Swift BAT transient monitoring programme spanning 20052016 May (Krimm, private communication). This supersedes the low-significance detections we tentatively reported in S11 . All of our sources, listed in Table 1 , have coverage in either the NRAO VLA Sky Survey (Condon etal. 1998 ) or Sydney University Molonglo Sky Survey (Mauch etal. 2003 ), three are also covered by the VLA Faint Images of the Radio Sky at Twenty centimeters survey (White etal. 1997 ) and one by the Westerbork Northern Sky Survey (WENSS; Rengelink etal. 1997 ). A search in radio catalogues at these positions results in no spatially coincident radio sources (within 30arcsec radius). However, XMMSL1J164859.4800507 (for which
Paper: 2016A&A...595A..53B (fulltext)
of one year of on-bulge time, but spread over several years of the mission. Table 5 lists the areal coverage and survey efficiency for a number of current and proposed ground and space microlensing surveys. For the space missions (Euclid and WFIRST) the efficiency in Table 5 is simply given by the likely total on-bulge observing time of the proposed programme. For the ground-based surveys the efficiency represents the fraction of the yearly rate likely to be achievable within a given observing season. The numbers in Table 5 can be used to scale the rate yields in Table 3 to approximate expected seasonal rates for a given survey in the limit of perfect detection capability across all event timescales. Figure 9 shows the maximum event rate detectable as a function of survey area for ground and space surveys listed in Table 5 . The actual area of the high-cadence regions of each survey is indicated on each line. We note that these maximum rates are obtained by rank sorting regions of the rate maps by the highest yielding regions and therefore a given total areal coverage may correspond to non-contiguous regions which do not correspond to the actual area being surveyed by the surveys. They instead represent the largest event yield which can be obtained for a survey with the same total areal coverage, wavelength and sensitivity. The advantage of a multi-site ground-based survey like KMTNet over single-site surveys is apparent from Fig. 9 . However a high-cadence near-IR survey with VISTA covering a similar area could prove even more effective owing to the increased sensitivity of near-IR observations. Figure 9 shows how space-based surveys are particularly effective for detecting low-mass exoplanets. Even though Euclid and WFIRST will cover areas of only around a couple of square degrees, their detection yield for Earth-mass FFPs is expected to be 12 orders of magnitude higher than ground-based survey seasonal rates which span an area of sky 10 times larger. For Jupiter-mass planets the space mission yields are 1030 times higher, and so are still likely to exceed yields from the ground, with the possible exception of a high-cadence VISTA survey operating over at least a decade. Overall, a short space-based campaign over a modest area can significantly outperform a gtound-based campaign and can provide important statistics on the FFP population over a range of FFP mass. In our simulation, the population of one FFP per star is assumed, but as Sumi et al. (2011) mentioned, the population of planets may be larger than stars. The detection probability of FFP is therefore expected to rise from our event rate values. Moreover, Varvoglis et al. (2012) and Wang et al. (2015) calculated the motion of FFPs in a cluster and resulted they tend to travel within the cluster unless any outside effect such as a close encounter with other clusters happens. The capture and release of FFPs ( Varvoglis et al. 2012 ) may also temporarily and slightly affect the
Paper: 2019MNRAS.485.3661F (fulltext)
time is dominated by mid latitudes where R ∼ r , so magnetic energy can only be dissipated at the pole as fast as it migrates from mid latitudes. Assuming loops of azimuthal field are totally dissipated near the pole, their effective damping rate is equal to their migration rate γ mig . Because the loop migration is essentially a random walk process, the migration time-scale is where N step is the migration length divided by a step length, and t step is the time it takes to complete each step. The number of steps is N step ∼ r /ξ ⊥ . The time of each step is a magnetic diffusion time across a radial wavelength, . In order for the the instability to grow, this diffusion rate must be smaller than the growth rate . So we have . Then using equations ( B3 ) and ( B4 ), we have This maximum migration rate will be realized when ω A ∼ ω c , with the critical field strength ω c defined by equation ( 1 ). When ω A ∼ ω c , field loops can reconnect with loops of opposite oscillation phase (i.e. those separated by radial distance ∼1/ k r ) after ∼1 oscillation cycle, such that they can migrate by a distance ∼ξ ⊥ each oscillation cycle. When ω A > ω c , reconnection requires many oscillation cycles, the migration rate will depend on the magnetic diffusivity, and it will be smaller than equation ( B5 ). Why do we still think the instability damping rate is δ v A / r ? Let’s consider whether migration of magnetic loops toward the pole can destroy them at faster rates. Assuming a loop reconnects with its neighbour after being displaced horizontally by ξ ⊥ at a rate , the two loops have moved apart from one another at a speed . Then the maximum rate at which the instability can be damped due to loops migrating to the pole is But since ω A Ω, γ diss is smaller than the damping rate δ v A / r . The actual destruction rate is likely much slower due to the random walk process discussed above, and is given by equation ( B5 ). This means that unstable perturbations will damp faster by weak turbulence than they will by migrating toward the pole. So the instability is still limited by weak turbulence, and setting the growth rate equal to the damping rate still implies The migration rate of equation ( B5 ) implies that the background field, whose energy density is , is destroyed at this rate. The assosiated energy damping rate per unit volume is The turbulent energy damping discussed in the paper will operate regardless of the loop migration, and will be more important when the reconnection time-scale is longer than an oscillation times scale. Hence, while damping from loop migration may be relevant when ω A ∼ ω c , we do not expect any of our scaling arguments or results to be altered. APPENDIX C: EFFECTIVE STRATIFICATION As discussed in Spruit ( 2002 ), the effective Brunt–Väisälä frequency N eff depends on the thermal diffusion time-scale across the Tayler instability length-scale. This thermal diffusion time-scale in turn depends on N eff . Spruit ( 2002 )
Paper: 2017ApJ...851L..27M (fulltext)
of the accretion disk (White Swank 1982 ). The spectral evolution during dips depends on the ionization state of the absorbing material (Diaz Trigo Boirin 2016 ). In the case of XMM0042, we do not detect any significant spectral variation possibly due to the low statistics. Van Paradijs McClintock ( 1994 ) study the optical emission from LMXB disks; if X-rays from the central source are reprocessed by the accretion disk, this implies that the optical luminosity in the V-band M V scales with the X-ray luminosity and size of the accretion disk. We considered the relation from Van Paradijs McClintock ( 1994 ) between L X , the orbital period, and M V , taking into account erg s −1 (following Van Paradijs McClintock 1994 ), our orbital period of 4.01 hr, and our mean luminosity erg s −1 :obtaining (1 σ error) for the optical counterpart. Among the 17 positionally consistent HST candidate optical/UV counterparts reported in Yukita et al. ( 2017 ), all but one are very similar to the surrounding stellar populations. The remaining source shows an excess in the blue band, making it an interesting potential counterpart of the accretion disk of an LMXB. From the blue apparent magnitude of the counterpart m B = 24.78, standard extinction toward M31, and Equation (4) from Barnard et al. ( 2012 ), we obtain an absolute visual magnitude . This is consistent (within 2 σ ) with the one derived from the X-ray luminosity and period of XMM0042, thus making it the likely optical counterpart. The X-ray spectrum of XMM0042 (as seen by Yukita et al. 2017 and in this work) is quite hard, with and . This is consistent with what is seen in dipping LMXBs in our Galaxy, where a wide range of spectral properties is displayed, with photon indexes varying from 0.4 to 2 and cutoff energies from 3.5 to 80 keV (Diaz Trigo et al. 2006 ). The most luminous Galactic dipper, X 1624–490, has an X-ray luminosity in the 1–30 keV energy band of erg s −1 (in the 0.5–50 keV energy range—Galactic dippers reach persistent luminosities only ∼3 times lower than XMM0042; Balucinska-Church et al. 2000 ; Iaria et al. 2007 ) and a much softer spectrum. The period, optical counterpart, and the spectrum of XMM0042 are reminiscent of a dipping LMXB system. The mean luminosity we found through our XMM-Newton analysis, ranging from 0.8 to erg s −1 , makes XMM0042 the most luminous dipper to date. Trudolyubov et al. ( 2002 ) found the first dipping source of M31, located in Bo 158. The period of the dipping behavior was 2.78 hr with a flux modulation of ∼83%. A study of the other XMM-Newton , Chandra , and ROSAT observations (Trudolyubov Priedhorsky 2004 ) revealed the amplitude of the dips to be anti-correlated with the source luminosity ( erg s −1 ), disappearing at high luminosities. The source also showed hour-timescale and month-timescale luminosity variations. This source has a rather hard spectrum, but it has never been observed by NuSTAR , so we have no information about the cutoff energy or