adsabs/WIESP2022-NER huggingface dataset:
Paper: 2016MNRAS.461.3075F (fulltext)
found that Centaurs can produce HTCs, in particular they are trapped very often in the 1:6 MMR with Jupiter (energy x ≃ 0.058 au −1 ) which coincides with the clustering in the observed energy distribution of HTCs (Fig. 3 ). It is worth noting that our sample of observed Centaurs contains a few objects in retrograde orbits. To produce both the observed HTCs and JFCs, a flattened source is required but it should contain a small fraction of high-inclination and retrograde comets. A near-ecliptic, prograde population of Centaurs ( i ≲ 25°) will essentially lead to a near-Earth population of bodies with Tisserand parameters T J > 2, typical of JFCs (Levison Duncan 1997 ). It is not the purpose of this paper to discuss the ultimate origin of Centaurs, though we can say a few words. A flat source in the outer planetary region such as the Scattered Disk will essentially produce a population of Centaurs in prograde, low-inlination orbits (Di Sisto Brunini 2007 ), though it may be possible that some SDOs scattered outwards attain Oort cloud distances where galactic tides switch their orbits from prograde to retrograde (Levison et al. 2006 ). Other possibility is that some Centaurs come from the capture of Oort cloud comets in the outer planetary region as discussed by Emel'yanenko Bailey ( 1998 ) and Emel'yanenko, Asher Bailey ( 2005 ). In the end, some HTCs may come from the Oort cloud but through a different dynamical route that direct insertion of Oort cloud comets in the near-Earth or the Jupiter region. In order to check the results found for the simulated HTCs produced from the evolution of the Centaur sample, we also studied the dynamical evolution for 1 Myr to the past and to the future of 15 observed HTCs with q 1.3 au and energies 0.04 x 0.08 au −1 (orbital periods 44.2 P 125 yr). We confirm that HTCs do not diffuse randomly in a but tend to be captured in the 1:6 exterior resonance with Jupiter generating a concentration at x ∼ 0.058 au −1 . This peak is clearly shown in Fig. 17 which is a histogram of frequencies of the average 〈 a 〉 over periods of 500 yr. Figure 17. Frequency distribution of the time-average 〈 a 〉 (over 500 yr) for the observed sample of 15 HTCs with q 1.3 au and energies 0.04 x 0.08 au −1 (periods 44.2 P 125 yr) integrated over 10 6 yr in the past and in the future. Other remarkable feature is that even though we obtain about 4.7 per cent of LPCs that end up colliding with the Sun for a purely dynamical evolution, this percentage is practically null when we consider appropriate physical lifetimes ( N phys ( q = 1) ≲ 500 revolutions). The reason is that LPCs approach the Sun in a smooth way through a multi-step process, that takes hundreds to thousands revolutions, so it is very likely that such comets will disintegrate well before reaching sungrazing states. The inclination distribution of the old LPCs shows clear departures from randomness that can be described as a slight predominance of prograde orbits that is
Paper: 2019ApJ...881...48K (fulltext)
high count rates occasionally occurred in the so-called tentacle region (Forster et al. 2014 ) near the SAA. Thus, we ran nupipeline to reject times with high count rates by setting options saamode=optimized and tentacle=yes . Background rates during the J0822+2241 observation were stable and low, and thus we adopted saamode=none and tentacle=no . We defined source regions as 30′′-radius circles centered at each optical position by taking account of the FWHM of the NuSTAR point-spread function (PSF; ≈18′′). The size is much larger than the typical size of GPs (1′′; Cardamone et al. 2009 ). Background regions were off-source circular regions with a 30′′-radius on the same detector. Then, we produced source and background spectra, and response files using the nuproducts task. The products of FPMA and FPMB were combined to provide better statistics by using the addascaspec command. The systematic uncertainty between the two modules is likely much smaller than the statistical uncertainty. Figure 3 shows the obtained spectra, as well as the background contribution. This clearly illustrates that we detected no significant emission from J0749+3337 and J0822+2241 even in the most sensitive 8–24 keV band. Count rate upper limits at 3 σ in the energy range are 5.7 × 10 −4 counts s −1 and 4.5 × 10 −4 counts s −1 , respectively. They are converted into 2–10 keV luminosities of ≈2 × 10 43 erg s −1 and ≈1 × 10 43 erg s −1 , respectively, by adopting a power-law model with Γ = 1.7. In Section 5.2 , we further investigate how large absorbing column densities are needed to be consistent with the nondetection if there are AGNs with (Section 2 ). Figure 3. NuSTAR 3–50 keV and XMM-Newton 0.2–9 keV spectra taken from the source (black crosses with circles) and background (blue dashed crosses with diamonds) regions. The figures suggest that the GPs are detected in the soft band, but not in the hard band. 4. <italic>XMM-Newton</italic> Soft X-Ray Data Analysis The XMM-Newton data for J0749+3337 and J0822+2241 were obtained through the European Photon Imaging Camera/MOS (1 and 2) and PN detectors with a duration time of ≈44 ks and ≈34 ks. All observations were performed by adopting the Prime Full Window mode and the thin filter. We reduced the data following the XMM-Newton ABC guide. 13 13 https://heasarc.gsfc.nasa.gov/docs/xmm/abc/ The raw PN and MOS data were reprocessed using pipelines of epchain and emchain , respectively. To filter periods with a high background, we created PN background light curves in the 10–12 keV band with PATTERN=0 (single events), and those of the MOS in energies above 10 keV with the same pattern selection. Regarding the J0822+2241 data, we adopted background count rate thresholds of 0.35 s −1 and 0.40 counts s −1 for the MOS and PN cameras, respectively. These are the recommended values in the guide. In contrast, we found two high-background flares during the J0749+3337 observation, and could not clearly remove the tails of the flares
Paper: 2018ApJ...855..132F (fulltext)
we study (i) a constant 29% contamination and (ii) the stellar mass and redshift-dependent [N ii ]/H α flux ratios predicted by our model (Equation ( 3 )). To derive the intrinsic H α LF using our model [N ii ] contamination (Φ model ), we start with the observed ([N ii ]+H α ) tot LF (Φ obs ( L )) measured by Colbert et al. ( 2013 ) at 0.9 z 1.5. We obtain this LF from their published H α LF (, , and α = −1.43) by dividing the luminosities by a factor of (1–0.29) to undo the constant total [N ii ] flux contamination fraction correction of 29%, which the authors applied. In the following, we treat Φ obs as the true observed LF. Importantly, this LF is not corrected for [N ii ] contamination and dust. Because of the evolution of the star-forming main sequence, Φ obs is redshift-dependent, but here we do not model this dependence across z = 1–2, as we are only interested in the effects of [N ii ] contamination and not the absolute number of galaxies. On the other hand, the stellar mass and redshift-dependent [N ii ] contamination correction will change the intrinsic H α LF across the redshift range studied here. To obtain Φ model , we choose an approach that only uses the measured ([N ii ]+H α ) tot luminosities as input and assumes the most likely underlying stellar mass distribution (robustly determined from other studies), from which we obtain the [N ii ] contamination from our model. This approach has the advantage that it enables an easy implementation and propagation of a variety of uncertainties into the final results. Furthermore, this method results in reliable intrinsic H α LFs even if the mass distribution is poorly measured due to the lack of sufficient multiwavelength data, as long as the selection function of the galaxy sample is known. Here we make use of the Schreiber et al. ( 2015 ) parameterization of the star-forming main sequence to derive the underlying stellar mass distribution (we comment below on possible shortcomings). We note that the choice of different parameterizations (e.g., Tomczak et al. 2016 ) should not change the following results. We start with a distribution of ([N ii ]+H α ) tot , which we sample from Φ obs . To obtain stellar masses for these galaxies, we use a “backward-engineering” technique. First, we convert the SFRs of the Schreiber et al. ( 2015 ) parameterization into H α luminosities using the Kennicutt ( 1998 ) prescription. Thereby, we include a dispersion of 0.3 dex measured on the SFR–versus–stellar mass main sequence. Second, we redden the H α luminosities according to the relation between A H α (H α extinction) and stellar mass robustly derived from the spectra of local galaxy samples in SDSS (Garn Best 2010 ). This relation holds for the WISPS sample at 0.8 z 1.5, as shown in Domínguez et al. ( 2013 ). Third, we add the contribution of [N ii ] to the H α luminosity by using our model. Finally, this translation between dust-reddened ([N ii ]+H α ) tot luminosities and stellar masses allows us
Paper: 2016MNRAS.455.3674V (fulltext)
may be higher given that an unmeasurable component of light is lost during this period. No conclusion can or should be drawn from this point, which is presented only for the sake of completeness. 3.2 He <sc>i</sc> line 3.2.1 Equivalent width The equivalent widths of the He i line around the 2014 periastron are shown in Fig. 4 (b). The closed blue circles denote the equivalent width as measured by integrating from the base of the line, while the open green squares denote the measurement obtained from the Gaussian fit, relative to the continuum. Shown for comparison are the measurements around 2010 reported by Chernyakova et al. ( 2014 ) (red crosses). As with the Hα line, the He i line shows a general increase in the absolute equivalent width which peaks after periastron, though the growth is not as great. The variation in the strength of the line is consistent with the variability observed in the Hα line. 3.2.2 V/R In general, the violet component of the double peaked He i line is more dominant, as is seen in the equivalent width | W λ | of the violet component and the ratio of peaks of the violet to red (V/R) components. The V/R variation (Fig. 4 c) has been measured from the ratio of the heights of the Gaussian fits to the Violet and Red component of the He i line. The height is measured relative to the continuum and the error is given by the statistical error of the fit. Around the first disc plane crossing, there is a marked increase in the asymmetry of the line, while there is a decrease in V/R around periastron and a small increase in the red component around the second disc plane crossing. This clearly illustrates the asymmetry of the He i line and its variation around periastron. The V/R variation lies between 0.93 V/R 1.65 during the observed period. 4 DISCUSSION Around periastron the equivalent width of the Hα line smoothly increased to a maximum of −72.7 ± 0.7 around ∼τ + 13 d, after which it decreased to values consistent with the pre-periastron levels. The equivalent widths after ∼τ + 13 d are consistent with the previous periastron passage (Chernyakova et al. 2014 ) but appear to be consistently lower before this. This is clearly seen before periastron where the average equivalent width (and error) of the Hα line was −62.9 ± 3.1 Å in 2010, but −56.7 ± 0.6 Å in 2014. We interpret this as being due to variability in the Be star circumstellar disc which could well change from periastron to periastron. This may imply that the mass of the circumstellar disc was lower before the 2014 periastron passage (as compared to the 2010 periastron passage). Since the circumstellar disc influences the shock front around the pulsar, the differences in the circumstellar disc properties will be an important contributing factor to the variations in the multiwavelength emission observed around different periastra. There is also a marked change in the disc around the time of the first disc plane crossing (τ − 17.5 d), as is noted by the increase in the
Paper: 2017ApJ...835...42G (fulltext)
not been identified due to the lack of counterparts in the radio/X-ray or GeV γ -ray bands (Aharonian et al. 2005 , 2008a ). Therefore, searching for the counterparts of these unidentified VHE sources in longer-wavelength bands is very helpful to identify and classify them. In the GeV band, more than 7 years of survey data from the Fermi Large Area Telescope ( Fermi -LAT) provides a good opportunity to hunt for the unidentified VHE sources. Indeed, several unidentified VHE sources have been detected with the Fermi -LAT data (Hui et al. 2016 ). HESS J1427−608 is one of the unidentified VHE sources discovered by the HESS Galactic Plane Survey (Aharonian et al. 2008a ). The TeV image shows that it is only slightly extended with a symmetric Gaussian function with a width of σ = 3′. This source is detected with an integral flux above 1 TeV of F γ (>1 TeV) = 4.0 × 10 −12 erg cm −2 s −1 and a power-law spectrum with an index of Γ = 2.16. No nearby SNR or pulsar was detected in the vicinity of HESS J1427−608 (Green 2014 ). The X-ray counterpart of it, Suzaku J1427−6051, was reported by Fujinaga et al. ( 2013 ). The X-ray spatial morphology is clearly extended with a Gaussian width of σ = 09 ± 01, and the spectrum is well fitted by a power law with an index of Γ = , implying nonthermal X-ray continuum emission. The unabsorbed flux in the 2–10 keV band is . Fujinaga et al. ( 2013 ) have discussed the possible nature of HESS J1427−608 as a PWN or nonthermal SNR. Each one can explain some of the observational facts but face difficulties. No radio counterpart was detected in the direction of HESS J1427−608. In the second Fermi -LAT catalog (2FGL; Nolan et al. 2012 ), a GeV source, 2FGL J1427.6−6048c, which is located in the vicinity of HESS J1427−608, was reported. However, their central positions differ moderately from each other. Furthermore, the power-law spectrum index of 2FGL J1427.6−6048c is Γ = 2.7 (Nolan et al. 2012 ), which cannot connect with the TeV spectrum of HESS J1427−608 well. It is very likely that 2FGL J1427.6−6048c is not associated with HESS J1427−608. Meanwhile, it should be noted that no source in the third Fermi -LAT source catalog (3FGL; Acero et al. 2015 ) is associated with 2FGL J1427.6−6048c. In this paper, we carry out a complete analysis of this region using more than 7 years of Fermi -LAT Pass 8 data to investigate the γ -ray emission of HESS J1427−608. In Section 2 , we describe the data analysis and results, including the spatial, spectral, and timing analyses. The possible nature of HESS J1427−608 based on the multiwavelength spectral energy distribution (SED) fitting is discussed in Section 3 , and the conclusion of this work is presented in Section 4 . 2. DATA ANALYSIS 2.1. Data Reduction The following analysis is performed using the latest Fermi -LAT Pass 8 data with the Source event class from 2008 August 4 (Mission Elapsed Time 239557418) to 2015 November 4 (Mission Elapsed Time 468288004). Considering the impact of
Paper: 2017MNRAS.466.3513S (fulltext)
the dynamo action in astronomical bodies. We choose to work with periodic cuboid domains, because it allows us to study the dynamo properties of spatially periodic cellular flows. We vary the scale separation in the system through the reconstruction of the computational domain and the characteristic length-scale of the imposed velocity field. The first flow under consideration is the well-known ABC flow (named after Arnold, Beltrami Childress), recently reviewed by Galloway ( 2012 ). The second is a modulated version of the first, with imposed long-wavelength perturbation on the scale of the system. Such periodic flows are guaranteed to produce dynamo action for a wide range of finite values of R m (Roberts 1970 ). We study the kinematic dynamo action in domains of various aspect ratios, to examine the effect of changing the scale separation of the system as well as the effects of imposing a two-scale perturbed flow on the system in the kinematic regime. The first part (ABC flow) of the problem considered here has recently been investigated in a different way by Cameron Alexakis ( 2016 ), who carried out a theoretical study of kinematic dynamo using the Floquet formulation that was originally devised for the Mathieu equation (Whittaker Watson 1902 ). The authors adopted the concepts in Roberts ( 1970 ) to solve the induction equation in the presence of the ABC flow for arbitrary large-scale separations, showing a transition from a large- to small-scale dynamo through a change in the growth rate. The results from this analysis are in general agreement with the findings presented below as the scale separation becomes arbitrarily large. However, their paper does not include the results for Floquet waves with length-scales comparable to the size of the imposed flow, although Galanti, Sulem Pouquet ( 1992 ) noted that such eigenmodes play a crucial role in maintaining the dynamo for large R m . In this paper, we utilize a discrete wavenumber decomposition that is determined by the aspect ratio of the computational domain. This allows us to investigate the transition from a large- to small-scale dynamo with a special emphasis on the length-scale selection of the fastest growing eigenmode of magnetic field and focus on the comparison of the results from the classic ABC and the modulated perturbed flow cases. Our study is based on the results of numerical simulations using a pseudo-spectral method in a triply periodic domain and analytic estimates based on the scale-separation arguments. The choice of the imposed flows allows for the scenario wherein a large-scale dynamo would be obtained at the onset. We then vary magnetic diffusivity in order to study how the dynamo mechanism changes as we increase R m . There are two possibilities: (i) the large-scale structure is preserved due to either the large aspect ratio of the system or the imposed long-wavelength modulation; or (ii) the energy is passed from the large scales to the small scales as
Paper: 2018MNRAS.476.5453G (fulltext)
the result by about a factor of 2. Therefore, the shock width in terms of the pair unloaded Thomson optical depth is Fig. 6 shows , obtained from integration of the full equations (without assuming x l > 1), as a function μΓ u / f for several values of Γ u (i.e. on each curve, Γ u remains constant while f varies). It shows that equation ( 13 ) provides an order of magnitude estimate of the shock width as a function of the photon escape fraction. Fig. 7 shows the Lorentz factor profile as a function of . It shows how the shock becomes narrower when the fraction of escaping photons increases. Once the optical depth to infinity becomes too small and photon escape starts affecting the shock structure, the shock adjusts itself and becomes narrower to match the finite optical depth. This is in contrast to a Newtonian shock that cannot adjust itself, and therefore full shock breakout follows within a dynamical time after photon escape starts becoming non-negligible. Figs 8 – 9 show the mechanism that allows the shock to adjust its width – a rapid buildup of x l within a unit interval of the optical depth τ, which compensates for the lack of backscattered photons at τ > τ ∞ . As can be seen from Figs 7 – 9 , the structures of all shocks with different values of f converge while the flow is still relativistic, before it approaches the downstream (τ ≲ 1), even when photon escape strongly affects the shock width. This result supports our assumption that the conditions in the immediate downstream are similar in all shocks, even in the presence of a significant photon escape. Figure 6. The shock width as a function of μΓ u / f . On each of the four curves, f varies while Γ u = 5, 15, 50 and 150 remains constant. is measured using the full analytic solution with η = 0.5 and a = 2, and is defined as the pair unloaded Thompson optical depth, where Γ = 0.9Γ u . The circles mark the location where , at which equation ( 13 ) predicts a transition in the shock width dependence on f . The solid black line is a plot of , and is given for guidance. This figure indicates that the width is given by equation ( 13 ) to within an order of magnitude. Figure 7. Lorentz factor profiles as a function of for shocks with Γ u = 10 and different f values. Figure 8. Profiles of the loading parameter as a function of for shocks with Γ u = 10 and different f values. Figure 9. Profiles of the loading parameter as a function of τ for shocks with Γ u = 10 and different f values. 4 SHOCK BREAKOUT In the previous section, we derived the steady-state structure of an RRMS with photon escape. In astrophysical settings, the shock propagates in a medium having a finite Thomson depth to infinity . Note that we use to describe the optical depth of the circumstellar medium since no pairs are created before it is encountered by the shock. For typical settings, we expect both Γ sh = Γ u and to vary with time. Then, once , photons starts leaking out. However, since the shock width continuously
Paper: 2016MNRAS.455.2004Z (fulltext)
show evidence of hydrogen (i.e. longer durations). The fact that in 4U 1728–34 bursts with oscillations happen more often in the high/soft state, while PRE burst happen mostly in the low/hard state, suggests that the oscillations and the PRE phenomenon are both correlated with accretion rate and geometry, as proposed by Franco ( 2001 ), van Straaten et al. ( 2001 ) and Muno et al. ( 2004 ); however, the presence of bursts with oscillations in the low/hard state in 4U 1728–34 (and 4U 1636–53; Zhang et al. 2013 ) indicates that oscillations are not only driven by either mass accretion rate or whether the burst shows PRE or not, although mass accretion rate (and perhaps PRE) has an influence on the presence of oscillations. 4.2 X-ray spectral properties and burst oscillations We find a strong evidence for a link between burst oscillations and the spectral properties during the burst in 4U 1728–34, more specifically between the presence of burst oscillations and the changes in the apparent emission area during the decay of the burst. The left-hand panels of Figs 2(a) and (b) , and the discussion in Section 3.1 , show that oscillations in Type-I X-ray bursts in 4U 1728–34 are always associated with a blackbody radius that remains more or less constant for at least ∼2–8 s. A similar trend has also been found in the LMXB system 4U 1636–53 (Zhang et al. 2013 ). The changes of the blackbody radius could be due to either changes in the apparent emitting area of the NS during this phase, or to changes in the colour correction factor, f c . In the case of changes of the apparent emitting area of the NS, Spitkovsky et al. ( 2002 ) and Zhang et al. ( 2013 ) suggested that tail oscillations could be due to the spread of a cooling wake, which is formed by vortices during the cooling of the NS atmosphere. In this scenario, the speed of the cooling wake depends upon the latitude at which the burst ignite, since the speed of the cooling front near the equator is higher than that near the poles (Spitkovsky et al. 2002 ). If the bursts ignite at high latitude on the NS surface, during the burst decay time the front speed of the cooling wake is slow, the emission area changes slowly, and the asymmetric emission during the burst lasts long. These bursts would have oscillations and the blackbody radius remains constant for a while (see the left-hand panels of Fig. 2a ). If the bursts ignite at low latitude, during the decay time the front speed of the cooling wake is fast, the emission area, and hence the blackbody radius, changes fast, and the asymmetric emission is short lived. These bursts would not show oscillations and the blackbody radius evolves very fast. We note that this interpretation has not been calculated in detail. We also note that burst oscillations also appear in the rising phase of the bursts (e.g. left-hand panels of Fig. 2b ). The mechanism of burst oscillations might be different between the rising and decaying phase. The burst oscillations