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24 The extended Chern-Simons Hamiltonian in first quantisation, with N α particles of each type α, reads H CS = 1 2m α Nα jα=1 p jα + e A * α (r jα ) +e δA CS α (r jα ) + e a • α (r jα ) 2 ,(3.1) where we absorb the average value of the Chern-Simons potential (2.2) into an effective vector potential A * α (r) = A(r) + A CS α . This definition yields the effective magnetic field ∇ × A * α (r jα ) = B * α (r jα ) given in Eq. (2.4). In Fourier space, the fluctuations δA CS α (q) are transverse, similar to the gauge field itself, as given by Eq. (2.3). Here, we have δA CS α (q) = δA CS α (q)e ⊥ q = h e|q| β K αβ δρ β (q)e ⊥ q . | cond-mat.mes-hall |
The energy E p is assumed to be large relative to the hopping amplitude between iridium and oxygen, and also to the hopping between two oxygens. Therefore, the energy E p controls the perturbation theory used to compute the effective tight binding model parameters, which are expressed in powers of t/E p . Symmetry arguments provide strong constraints on the form of the tight-binding model that obtains from perturbation theory. For instance, since spin-orbit coupling is time-reversal invariant, the most general pseudospindependent hopping appearing in Eq. ( 2) is of the form t iα,jα ′ = t 0 ij δ αα ′ + iv ij • σ αα ′ ,(6) where v ij is a real vector and σ = (σ x , σ y , σ z ) are the three pseudospin Pauli σ matrices. Note that v ij must be odd under exchange of i and j in order to ensure Hermiticity of H 0 . | cond-mat.str-el |
The inset in Fig. 1(a) shows the m(H) behavior for the H-annealed CNT samples at T = 4 K, from which the coercive field is estimated to be ~100 Oe. the data yields transition temperatures of 910 K and 1120 K for the low-field and saturation moments, respectively. Thus, the Curie temperature is estimated to be T C =1000±100 K. In previous studies of ferromagnetic carbon [6,11,13], the temperature dependence was fit using a two-dimensional (2D) model that accounts for anisotropic spin waves in conjunction with the Ising model. In the present case, however, the mean-field model provides an adequate fit. Further information on the ferromagnetism was obtained by measuring the dependence of magnetization on the annealing temperature and on the number of CNT walls. | cond-mat.mes-hall |
Due to the operating frequency the measured susceptibility is composed of two parts. The first is due to the magnetic moments in the sample, and may be either para-or diamagnetic depending on the material studied. The second is due to the screening of an rf field via the normal skin effect in metals. This screening is a diamagnetic contribution and is a measure of changes in resistivity [16]. Radio frequency susceptibility data presented herein were collected in a TDR operating at 23 MHz mounted in a 4 He cryostat. The design is similar to that presented in Ref. | cond-mat.str-el |
Within the Born-Oppenheimer approximation 19 , the Schrödinger equation for the motion of a hydrogen atom with mass M is - h2 2M ∇ 2 ψ(r) + U (r)ψ(r) = Eψ(r). (1) The 3-dimensional potential energy surface (PES), U (r), contains the contributions to the total energy from all Pd atoms in the slab, all hydrogen atoms adsorbed on the surface, and the hydrogen atom in question at r. When only the position of this hydrogen atom is varied, U (r) becomes, in effect, the potential energy surface for its motion. The PES was mapped out by calculating the adsorption energy of the hydrogen atom placed at different positions over a 3V H defect. We formed the defect within a periodic 3×3 surface unit cell, which is large enough to prevent significant interaction with periodic images in neighboring cells. The potential energy surface was sampled according to the importance of each region: sample points were densely distributed near points of interest such as the fcc sites, the hcp sites, and along the pathway between them, while points far from these locations were sampled less densely. Altogether over 3500 energy points were evaluated, distributed within about 60 planes parallel to the surface and separated by 0.1 Å. | cond-mat.mtrl-sci |
13 Neutron inelastic scattering data published by Naberezhnov et al. on PMN offered the first dynamical evidence of PNR in the form of a prominent broadening of the transverse acoustic (TA) mode that coincides with the onset of strong diffuse scattering at 650 K, 14,15 roughly the same temperature (≈ 620 K) as that reported by Burns and Dacol in their optical study of PMN. This temperature, commonly known as the Burns temperature, and denoted by T d in the original paper, is widely viewed as that at which static PNR first appear. Likely condensing from a soft TO mode, distinctive butterflyshaped and ellipsoidal diffuse scattering intensity contours centered on (h00) and (hh0) Bragg peaks, respectively, are seen below T d in both PMN 16,17,18 and PZN. 19 Similar diffuse scattering patterns were subsequently observed in solid solutions of PMN and PZN with PbTiO 3 (PMN-xPT and PZN-xPT); however these patterns appear only on the Ti-poor (relaxor) side of the well-known morphotropic phase boundary (MPB). 20,21 The polar nature of the strong diffuse scattering, and thus its associ- 27 Middle -diffuse scattering intensity contours measured at (110) below (300 K) and just above (650 K) the Burns temperature T d using cold neutrons by Hiraka et al. | cond-mat.mtrl-sci |
A height of 28 nm corresponds to a normalized distance of z 0 /λ 0 ≈ 0.02, where we can reasonably expect the dipole approximation to hold (the spatial dependence will be shown below). Our material loss numbers (γ/ω 0 = 0.01 -0.001) are close to the state-of-the-art for metamaterials, but they are significantly greater than those used in some previous waveguide studies, where enhanced PFs were demonstrated with γ/ω 0 ≈ 10 -10 -10 -8 [27], or no loss at all [23]. For metamaterial applications, a useful figure of merit (FOM), is FOM = -Re(n)/Im(n), with a larger FOM indicating a less lossy metamaterial. The current FOM for typical metamaterial is of order 100 at GHz frequencies and drops to 0.5 at optical frequencies (380 THz) [25]. However, there are methods to improve these FOMs as they are not fundamental material properties; for example, Soukoulis et al. have suggested that the FOM can be improved by a factor of 5 at optical frequencies, and after optimizing their fishnet design, they have demonstrated that the FOM can be around 10 at 380 THz (cf. | cond-mat.mes-hall |
As suggested earlier, 21,45,46 the glass-like freezing of the magnetic moments for x 0.2 is most likely caused by a competition of ferromagnetic and antiferromagnetic exchange interactions, which can prevent a spontaneous symmetry breaking at a well-defined temperature. Hence, the entropy continuously decreases upon cooling. The existence of such a frustrated spin-glass phase may also affect the ferromagnetic order (or clusterglass phase) present for larger x. Probably, it is the reason for the absence of a clear anomaly in c p of the compounds which are located close to the boundary between spin-glass behavior and ferromagnetic order and also explains the rather broad transitions observed for the higher doped compounds. IV. La1-xCaxCoO3 The Ca substitution enhances the intrinsic bond-length mismatch responsible for the structural distortion, as its ionic radius is much smaller than that of La. | cond-mat.str-el |
This notion is often called Moore-Penrose inverse in literature, see, e.g. : R. Penrose, Proc. Cambridge Philos. Soc. 51, 406 (1955); A. Ben-Israel and T. N. E. Greville, Generalized Inverses: Theory and applications, Pure and Applied Mathematics (Wiley, New York, 1974), ISBN 0-471-06577-3. 29 Throughout the text, all sums over the component indices (α, β, . | cond-mat.mes-hall |
To this purpose we standardly replace the local part of the propagator of the conduction electrons by a constant ∆(ǫ) = π k |V k | 2 δ(ǫ -ǫ(k)) . = ∆ the value of which we set as the energy unit. With this simplification we expand all dynamical and thermodynamic quantities in powers of the interaction strength U and rearrange appropriately the resulting series. The principal idea of the parquet approach is to derive the self-energy and other thermodynamic and spectral properties via a two-particle vertex Γ. This vertex is standardly represented via Bethe-Salpeter equations. The Bethe-Salpeter equations express the full twoparticle vertex by means of irreducible ones that can be generically represented as 16Γ = Λ α + [Λ α GG] α ⋆ Γ (2a) where Λ α is the irreducible vertex in channel α, the brackets [. | cond-mat.str-el |
The result are characteristic van Hove singularities (VHS) in the density of states (DOS) [3]. Bulk (i.e. bundled) SWNT show an optical absorption peak at ∼ 4.5 eV due to transitions of the π electrons [4]. In vertically aligned SWNT (VA-SWNT) one finds the same peak position for onaxis polarization and an additional peak for perpendicular polarization at ∼ 5.2 eV [5]. Further information can be obtained from collective electronic excitations (plasmons) beyond the optical limit [6] (i.e. momentum transfer q > 0). | cond-mat.str-el |
The influence of a strong, localized electron-phonon interaction [9] can be simulated by the introduction of an effective negative-U Anderson model. In such system the charge Kondo effect leads to screening of charge fluctuations on the localized impurity by low-frequency pair fluctuations in the leads [10]. It has recently been shown that in the case of large electron-phonon coupling the original singleimpurity Anderson model, coupled to local phonon degrees of freedom, can be mapped onto an anisotropic Kondo model [11]. Increasing the electron-phonon coupling leads to a suppression of the conductance plateau width [11,9] as a function of the gate voltage. We study a double quantum dot (DQD) in a side-coupled configuration (figure 1), connected to a single conduction-electron channel. Similar systems have been studied using various analytical techniques such as the non-crossing approximation [12], embedding technique [13] or slave-boson mean field theory [14,15], and recently also using more accurate numerical renormalization group (NRG) calculations [16]. | cond-mat.str-el |
Obviously, if the incoming wave numbers have a range of ∆k, a∆k cannot be too large (i.e., comparable to unity), otherwise the interferences are smeared out. As we can see, there is a size effect here: at a given temperature T , the width of the equilibrium energy distribution is fixed, it has the order of k B T , and ∆k is also given by the dispersion relation. Thus if we can decrease a∆k e.g. by building smaller de-vices, then temperature related effects will become less pronounced. In other words, miniaturization, besides its obvious advantages, can help to avoid interferences to disappear (and also decrease dephasing caused by random scattering, simply by increasing the mean free path -system size ratio). 1)). | cond-mat.mes-hall |
(23) By our conjecture that Eq. ( 21) faithfully represents bosons in the LLL, the ground state of this Hamiltonian at filling factor 1/2 is then a featureless Mott insulator with topological order. We expect this to be true when K < V , where V is the strength of the nearest-neighbor repulsive interactions (U is the dominant interaction energy scale and can be taken to be large compared to V ). Assuming U ≫ K, V , the bosons can be taken to be hardcore, i.e. with double occupation of any site of the lattice prohibited. Using Holstein-Primakoff transformation 17 between hard-core bosons and spins of magnitude 1/2, we can rewrite Eq. | cond-mat.str-el |
The heat capacity and magnetoresistance measurements were performed using a Quantum Design physical properties measurement system. 151 Eu Mössbauer spectra were recorded at various temperatures using a conventional acceleration spectrometer with a 151 SmF 3 source. Laue diffraction spots were recorded on a Huber Laue diffractometer, while powder diffraction pattern was recorded using Phillips Pan-Analyitcal set-up. Electron probe micro-analysis (EPMA) measurements were performed on a CAMECA SX100 electron microprobe. III. RESULTS AND DISCUSSION In order to study the magnetic properties along the principal crystallographic directions, the single crystal of EuPtSi 3 was subjected to Laue diffraction; the good quality of the single crystals with four fold symmetry was confirmed by Laue diffraction spots. | cond-mat.str-el |
This force can then damp the thermal, Brownian motion and cool the oscillator by coupling to the effectively "colder" bath of the measurement. In this way, the fundamental limit of this cooling technique is the sensitivity of the measurement. In fact, it is only a Heisenberg-limited measurement that could completely cool to the quantum ground state [1]. So while this technique has been used to greatly reduce the effective temperature of a mechanical mode [2,3,4], true ground state cooling is not a realistic possibility until a quantum limited measurement has been achieved. While mesoscopic measurement techniques that can couple to lighter and smaller resonators have produced the closest approach to this measurement limit [5,6], technical issues have prevented measuring at the quantum limits of displacement sensitivity. In contrast, the most precise measurements in term of absolute displacement sensitivity use interferometric techniques with visible light to infer the mechanical position [2]. | cond-mat.mes-hall |
Therefore, it was concluded that a source of the excess charge is the antidot, and its conductance is mediated by Coulomb charging. Additional evidence for the CB effect follows from measurement of the conductance as functions of both magnetic field and source-drain bias, where clear and regular Coulomb diamonds were observed. 10 The doubled frequency conductance oscillations measured in Refs. 4,8 are also a strong indication of the Coulomb charging in the antidot. Thus, we conclude that 1/f c periodicity observed by Goldman et al. 14 might be a result of the CB effect. | cond-mat.mes-hall |
The resulting initial DW configuration at time t = 0 (when the spin current is switched on) with arbitrary center x 0 is then given by n(x, 0) = cosh -1 x -x 0 w êx + tanh x -x 0 w êz ,(34) where w = J/K is the initial width of the wall. We then determine the time evolution of this configuration from the LLG equation in the presence of a spin polarized current J s switched on at t = 0, taking the full spin torque T = T 1 + T 2 , see Eqs. ( 25) and (32). Due to the complexity of the resulting LLG equation, we have to solve it numerically. To that end, we spatially discretize the LLG equation into a discrete 1D spin chain, with spins n i at x = ia (integer i) for lattice constant a. The resulting first-order ordinary differential equation (in time) is then solved numerically starting from the initial configuration (34). | cond-mat.mes-hall |
Also, there is a decrease of the overall density of π * states in the N(1s) spectra, which is a general trend found in the FL-CN x . The latter is made more evident from the spectra shown in Figure 11 displayed in Figure 12, showing the evolution of the π * /σ * intensity with growth temperature. We must recall that, in this case, the window for FL formation occurs at high growth temperature. Therefore, Fig. 12 corroborates that there is a clear correspondence between the formation of the FL structure and the decrease in the density of π * states occurring at the N atoms. Another striking results is that the density of π * states at the C atoms remains essentially constant, i.e. | cond-mat.mtrl-sci |
Our functional was then applied to parabolically confined quantum dots by varying the number of electrons N in the dot and the confinement strength ω. From the fully self-consistent DFT results summarized in Fig. 1 it is clear that our functional outperforms the LDA for the whole range of N and ω. A more quantitative picture can be obtained from Tables I and II where we present numerical values for the exchange energy obtained with different approximations, for both parabolically confined and hard-wall square 32 quantum dots, respectively. It is clear that our functional yields errors that are smaller by at least a factor of 4 than the errors of the simple LDA. We tested the performance of our GGA functional also in the large-N limit. | cond-mat.str-el |
Instead they result from the particular phonon structure in nanowires. Thus, the observation of these current replicas manifests the interplay between electron and phonon confinement in the system. To model the phonon scattering environment in the nanowire, we calculate the one-dimensional phonon modes of the nanowire within an isotropic elastic continuum model [4]. Our InAs nanowires typically have nearly perfect wurtzite crystal structure with the wire oriented along the 0001 direction. Neighboring stacking layers of the wurtzite structure agree with those in a zinc-blende structure in 111 direction [16]. Since material parameters for the InAs wurtzite lattice are largely missing, we take them from a rotated bulk zinc-blende structure instead. | cond-mat.mes-hall |
In the language employed in this article, one then finds 41 K ν (t) = 1 (2πh) 2 dA|t|P γ (A; |t|),(115) where the integral extends over all phase space points A and P γ (A; |t|) is the probability that A is on a periodic trajectory with period |t|. In the presence of timereversal symmetry K ν (t) is a factor two larger. Setting P γ = Ω -1 Eq. ( 115) reproduces the form factor of random matrix theory; 60 Taking P γ (A; |t|) to be the return probability in a diffusive medium one recovers the spectral form factor of a disordered metal. 41,57 Since no trajectories with small-angle encounters are involved in the semiclassical calculation of K ν (t), the leading contribution to K ν (t) has no Ehrenfest-time dependence. (This is different for quantum corrections to K ν that are of higher order in h. 61,62 Such quantum corrections are not considered here.) | cond-mat.mes-hall |
The one-electron density of LiH is plotted in the top graph of Fig. 3, where the lithium (hydrogen) atom is situated at negative (positive) x-positions, cf. the ion potentials v d (x). In all considered cases, the density shows a clear minimum between the nuclei, and correlations mainly alter the density around the hydrogen atom. In particular, we highlight, that the second Born ground-state density is in surprisingly good agreement with the exact result. In order to identify the intra-molecular electronic structure more closely, we, in addition, have computed the most relevant natural orbitals (NO) for the 1D LiH molecule, see Fig. | cond-mat.str-el |
Figure 7 depicts the self-consistently calculated Hall resistances, considering different modulation amplitudes V 0 for a fixed sample width (2d = 3 µm) and m p = 5. We observe that, the plateaus become wider from the high B field side, when V 0 is increased, i.e mobility is reduced. Such a behavior is now consistent with the experimental findings. Since the QHPs occur whenever an incompressible strip is formed (somewhere) in the sample and the modulation forces the 2DES to form an incompressible strip at a higher magnetic field, therefore the plateau is also extended up to higher field compared with the (approximately) non-modulated calculation, V 0 /E 0 F < 0.1. Our investigation of the impurities lead us to conclude that, one has to define mobility at high magnetic fields also taking into account screening effects in general and furthermore also the geometric properties of the sample such as the width and depletion length. As an example if we consider an impurity concentration of ≈ %1 the long range part of the potential fluctuation can be approximated to 900 nm. | cond-mat.mes-hall |
vertex correction. The explicit form of these diagrams is given by equations ( 9) and ( 19). The wavy line represents the Coulomb interaction, Eq. ( 3). The momentum sums are performed ask = d 2 k/(2π) 2 . Next, we calculate the bubble dressing due to the electron-electron interactions to first order in α. | cond-mat.str-el |
This expansion may be inserted into Eqs. (42, 43) and the integration performed approximately analytically. It is assumed that only modes with an energy ǫ q < k B T contribute appreciably to the integral. In performing the integration one has to distinguish two cases: If one is within the NAF sector the second order coefficient a 2 is nonzero. If one is at the boundary to the CAF regime a 2 = 0 and the dispersion is determined by the mixed fourth oder coefficient a ′ 4 . Performing the integration in this (classical) limit k B T ≫ ∆ h one obtains the approximate expressions Γ mc Γ 0 mc ≃ h c k B T ln k B T ∆ h (63) for 2J 2 < J 1 (NAF) Γ mc Γ 0 mc ≃ 2 h c k B T SJ 2 ∆ h 1 2 ln k B T (SJ 2 ∆ h )1 2(64)for 2J 2 = J 1 (NAF/CAF) In the corresponding quantum limit k B T ≪ ∆ h , both the heat capacity and the rate of change of magnetization with temperature have an activated behavior. | cond-mat.str-el |
The Hamiltonian of the tt ′ -t ′′ -J model on the square Cu lattice has the form: H t-J = -t ij σ c † iσ c jσ -t ′ ij ′ σ c † iσ c j ′ σ -t ′′ ij ′′ σ c † iσ c j ′′ σ + J ij σ S i S j - 1 4 N i N j . (4) Here, c † iσ is the creation operator for an electron with spin σ (σ =↑, ↓) at site i of the square lattice, ij indicates 1st-, ij ′ 2nd-, and ij ′′ 3rd-nearest neighbor sites. The spin operator is S i = 1 2 c † iα σ αβ c iβ , and N i = σ c † iσ c iσ is the number density operator. In addition to the Hamiltonian (1) there is the constraint of no double occupancy, which accounts for strong electron correlations. The values of the parameters of the Hamiltonian (4) for cuprates are known from neutron scattering, Raman spectroscopy, and ab-initio calculations. For La 2 CuO 4 the values are 9,10,11 : J ≈ 140 meV → 1 , t ≈ 450 meV , t ′ ≈ -70 meV , t ′′ ≈ 35 meV . | cond-mat.str-el |
In all cases, V = 2.0, β = 0.001, and t ′ , V , V ′ , β, and ω are given in units of t. tegrable chains with obc's only reduces the height of the lowest frequency peak in Re [σ(ω)] and, with increasing t ′ , its weight moves toward higher frequencies. An apparent difference between Re [σ(ω)] in systems with pbc's and systems with obc's (Fig. 3) is that, while the former exhibit a finite dc conductivity, the latter exhibit a vanishing one. The vanishing of the dc conductivity for finite systems with obc's is understandable by using an analogy with pbc's with a small but finite momentum q. For q = 0, we know from the conservation law [ω n(q, ω) ≃ q Ĵ(q, ω)] that the current must vanish in the dc limit. 7 One expects, in this case, to see diffusion, i.e., σ(q, ω) = σ(0, ω)/[1 -iDq 2 /ω] with q ∼ π/L for obc's. | cond-mat.str-el |
From the results for the local currents, we see that the currents take finite values on all sites, which actually happens immediately after applying the potential. This clearly distinguishes the breakdown mechanism induced by a linear profile from other spatial forms of the bias voltage. For instance, in the simplest case in which V i = 0 in the interacting region and V i = ±V /2 in the lead, the physics underlying the breakdown is quite different as we have verified in additional simulations (results not shown here). In this case, the redistribution of the charge inside the interacting region can be described as an effective doping of the MI region from the two interfaces. This implies that the bulk of the interacting region will experience the effects caused by turning on bias with a delay, set by the length of the interacting region. Density and current profiles Turning back to Fig. | cond-mat.str-el |
Based on symmetry arguments, we are able to conclude that the local conserved quantities impose very strict constraints on spin-spin correlations [11], and an extremely characteristic pattern emerges in the spatial distribution of spin-spin correlations. More interestingly, this pattern is protected against small size, open boundary, and thermal effects. It is also robust against small perturbing interactions that may be present in realistic experimental setups. Our main results are summarized in Fig. 2 where the characteristic ordered emergent correlation pattern of the Kitaev model are compared with the messy results of the anisotropic Heisenberg model shown in Fig. 3. | cond-mat.str-el |
(1.26) [57]. Although this wave function contains a very large number of free parameters, its energy is always higher than that obtained from the pBCS state, showing the importance of having long-range valence bonds. A further drawback of this approach is that it is not possible to perform calculations on large system sizes, the upper limit being N ∼ 40. The second RVB state is obtained by considering long-range valence bonds, with their amplitudes given by Eq. (1.26) and optimized by using the master-equation scheme [58]. While this wave function is almost exact in the weakly frustrated regime, its accuracy deteriorates on raising the frustrating interaction, and for J 2 /J 1 > 0.425 the minus-sign problem precludes the possibility of reliable results. | cond-mat.str-el |
For a constant number density of particles, the density of states at the Fermi surface, g, rises with increasing number of flavors as g ∝ ν √ E F ∝ ν 2/3 . Therefore, the screening length estimated with the Thomas-Fermi approximation [24] is κ -1 = (4πe 2 g) -1/2 ∝ ν -1/3 , and the ratio of the screening to Fermi momentum length-scale varies with number of flavors as p F /κ ∝ ν -2/3 . In the many-flavor limit ν ≫ 1, the screening length is much smaller than the inverse Fermi momentum, κ -1 ≪ p -1 F , and so the dominant electron-electron interactions have characteristic wave vectors which obey q ≫ p F . This is in direct contrast to the random phase approximation (RPA) where p F ≫ κ, although in both the many-flavor and the RPA, the same Green function contributions with empty electron loops dominate diagrammatically [2,3]. These diagrams contain the greatest number of different flavors of electrons, and as ν ≫ 1 therefore have the largest matrix element. Since q ≫ p F , the typical length-scales of the MFEG are short, this indicates that a local density approximation (LDA) could be applied. | cond-mat.str-el |
Thus, in the electron basis we get the same two-body interaction plus a one body potential, which is attractive if the two-body potential is repulsive. From now on we focus on the specific short-range (hard-core) interaction that corresponds to Haldane pseudopotential V m = δ 1,m for H h . After diagonalizing the Hamiltonian H h with N = 20 electrons in N orb = 28, we obtain the energy spectrum in Fig. 10. It is worth pointing out that since there is no edge confinement other than the momentum cut-off due to the choice of N orb , we have N h = N orb -N = 8 holes in the system. Not surprisingly, the largest angular momentum of the degenerate ground states is M 0 = 294, as expected from Eq. | cond-mat.mes-hall |
Beyond the inherent importance of exploring a complex phase of quantum matter, these systems are of great interest for device applications involving quantum computing 7 and photonics, 8 the latter due to nonlinear electron-phonon interaction effects. Here we present an infrared spectroscopic study of Bi 2 Se 3 in magnetic field. The original motivation for this work was to search for signatures of the topological insulator state, a proposal with twofold justification. First, the topological nature of such a material is theorized to be sensitive to the application of electric and magnetic fields, 9,10 and such a tuning of the surface states may display a spectroscopic signature. Second, the magnetic field allows the probing of band dispersion via the cyclotron resonance. As mentioned above, a defining characteristic of the topological surface state is the existence of an odd number of Dirac cones at the Fermi surface, 3 which in turn prescribes the presence of mass-less Dirac fermions moving at the Fermi velocity. | cond-mat.str-el |
Since the Volterra solution is odd, the n = 1 term of the stress at each dislocation vanishes as we sum over the neighboring dislocations on either side. The first nonvanishing term in equation 7 is the n = 2 term which has three contributions. (1) The n > 1 terms are the multiple expansions of the stress field 23 as well as nonlinear terms. (2) The nonlinear term in strain field has the form du/dx * du/dx, giving a power law of 1/r 2 . (3) Geometrical restrictions cause some grain boundaries to have flaws unequally spaced in the y-direction (though for the results given in figure 14 we have only explored geometries with equally spaced flaws). The grain boundaries geometries used in figure 14 do have flaws that are not aligned perfectly in the x-direction. | cond-mat.mtrl-sci |
At 60 K it is easier to magnetize the Mn along the (020) direction than along (100) p-cub axis (see figure 3a). While at the lowest measured temperatures, where the Gd influences the magnetization, the easier axis corresponds to the (100) p-cub axis(see figures 3b and 3c). Although more studies are necessary to clarify this point, the anisotropy difference of the Mn and Gd sublattices could produce canting between them. Figure 4 shows the saturation magnetization (M s ) obtained from hysteresis magnetic loops like those shown in figure 3. In the M s estimated a paramagnetic signal was subtracted. The possible phase coexistence is supported by the low M s value at 5 K 80 emu / cm 3 . | cond-mat.mtrl-sci |
The self-energy corrections are naturally absorbed into the renormalization of the Fermi velocity. The non-RPA vertex diagram at lowest order of perturbation theory was found to decrease the effective charge, meaning that in principle correlation effects at higher order must also be taken into account. Thus the ultimate asymptotic behavior of the static polarization function for α ∼ 1 remains an open problem. Acknowledgments We are grateful to D. K. Campbell, R. Shankar, V. Pereira, O. Sushkov, and A. Polkovnikov for many in-sightful discussions. B.U. acknowledges CNPq, Brazil, for support under grant No.201007/2005-3. | cond-mat.str-el |
( 1) that, in turn, disagrees with the experiment of Goldman et al. 14 The Hartree approach is known to describe well the electrostatics of the system at hand. This is confirmed by the good agreement with numerous experiments including, for example, formation of compressible/incompressible strips in quantum wires 28 and the statistics of conductance oscillations in open quantum dots. 25 Thus, the validity of the energy level evolution as well as the electron number oscillation presented in Fig. 5 is indeed qualitatively correct. The conductance, however, may be incorrect. | cond-mat.mes-hall |
Thus, the good correspondence of the calculated and experimental results shown in Fig. 4(d) is remarkable. The change in lineshape is marginal for U ≥ 2 eV. These results clearly indicate that the electron correlation is strong among Ir 5d electrons. U seem to be comparable or higher than the estimations for 4d transition metal oxides. [26,27,28] Such anomaly may be attributed to the narrow dispersion of the Ir 5d bands leading to significant local character in the corresponding electronic states. | cond-mat.str-el |
However, we already questioned the The transmission coefficient and its spin polarization for a GaAs/(AlAs) 8 /GaAs single barrier structure Figure 3 The transmission coefficient and its spin polarization for a GaAs/(AlAs) 8 /GaAs single barrier structure. The red solid line is for 'spin-up' and the blue dashed line is for 'spin-down'. Two cases of k || are shown. One energy window of large spin polarization exists in the vicinity of -0.2 eV for k || = 0.03 2π/a. No such window is obtained for k || = 0.06 2π/a. This shows that efficient spin polarization can be obtained independently of the magnitude of Dresselhaus splitting. | cond-mat.mtrl-sci |
( 12) automatically preserves proper uniform scaling of T s (see Ref. 31): T s [n γ ] = γ 2 T s [n] , n γ (r) ≡ γ 3 n(γr) . (14) Constraints that must be satisfied by the enhancement factors associated with any satisfactory GGA KE functional includet θ ([n]; r) ≥ 0,(15)as well as 28,32,33 v θ ([n]; r) = δT θ [n]/δn(r) ≥ 0 , ∀ r . (16) The quantity v θ is known as the Pauli potential. Constraint Eq. ( 15) implies the non-negativity of the GGA enhancement factor, F θ (s(r)) ≥ 0. | cond-mat.mtrl-sci |
( 10) and ( 11) using the quantum numbers that lead to the minimum energy, as explained above, analytical expressions are obtained for the ground state energy and momentum as a function of flux for L sites and N particles, E g (L, N, φ) and K g (L, N, φ). Defining φ = φ for odd N, φ = φ -π for even N,(14)and writing φ in the form φ = nint N φ 2π 2π N + φ r ,(15) where nint[x] denotes the nearest integer to x, one obtains E g (L, N, φ) = -2t sin(N π/L) cos( φ r /L) sin(π/L) ,(16)K g (L, N, φ) = N φ/L. (17) For odd N , the transmittance vanishes at φ = π due to the reflection symmetry of the system. 32 This argument does not work for even N because the ground state of H ring has orbital degeneracy for N + 1 electrons at φ = π. It cannot be applied either for M = L/2 (where the reflection symmetry is lost 8 ) or if the model includes hopping at large distances (as in Ref. 31). | cond-mat.str-el |
1B) the shape anisotropy produces the same effect. When the easy-plane anisotropy energy is much larger then all other energies, the deviations of n(t) from the in-plane direction are very small. An approximation based on such smallness is possible and provides an effective description of the magnetic dynamics in terms of the direction of the projection of n(t) on the easy plane, i.e. in terms of one azimuthal angle. In this paper we derive the equations for effective in-plane motion in the presence of spin-transfer effect and discuss their use by considering several examples. In the absence of spin-transfer effects the large easyplane anisotropy creates a regime of overdamped motion even for the small values of Gilbert damping constant α ≪ 1. | cond-mat.mtrl-sci |
No further assumptions were made either in the thickness or concentration profiles describing the corrosion layer. Results Since the clean spot can be considered a homogeneous sample, the 3 MeV proton PIXE spectrum can be used with GUPIX [11] to obtain an independent characterisation of this sample in order to validate the DataFurnace results. The Table 1 shows a good agreement between the DataFurnace and GUPIX characterisations of the clean spot (note that for DataFurnace all the data was considered as opposed to just the single PIXE spectrum used for GUPIX). The good fit for the RBS spectrum of the same sample (shown in Fig. 1) further confirms the validity of the DataFurnace analysis. Figure 2 shows the composition profiles for the elements present in the dark spot of the photographic plate. | cond-mat.mtrl-sci |
Very importantly, there is a large temperature drop, ∆T , at the nanoparticle-liquid interface which is a manifestation of the interfacial thermal resistance. Such resistance is caused by the mismatch of thermal properties between the solid and liquid components, and is also affected by the strength of the interfacial bonding. The interfacial thermal conductance, G, can be quantified via the relationship:jQ = G∆T [ 2 ] Where jQ is the heat flux across the interface and ∆T is the discontinuous temperature at the interface (see temperature profiles in Fig. 1). Fig. 2 (top panel) shows the relationship between the heating power P and the temperature drop at the octane-liquid interface. | cond-mat.mtrl-sci |
In Fig. 3 the temperature dependent magnetization (M) divided by applied field (H) reveals the primary difference between the Fe column members of this family and the Co column mem- compounds, the 1.85 K magnetization isotherms, measured with the applied field along [100], [110], [111] crystallographic directions, were found to be isotropic to within less than 5 %. This magnetic isotropy is not unexpected in the Gd-based intermetallics, in which the magnetism is mainly due to the pure spin contribution of the 4f shell of Gd 3+ . For T = Fe, Ru and Os the magnetization is representative of a FM-ordered state with a rapid rise and saturation of the ordered moment in a field of the order of the estimated demagnetizing field (magnetic domain wall pinning being low in these single crystalline samples). For T = Co, Rh and Ir the field dependent magnetization data are consistent with AFM-ordered states that can be field stabilized to fully saturated states in large enough applied fields. This fully saturated state is observed for GdCo 2 Zn 20 associated with a spin-flop transition (see discussion below). | cond-mat.str-el |
4 the evolution of Ω R vs H 0 . The large dispersion observed on Ω R at given H 0 definitively confirms that it is, as expected, a directional effect. All the measured Rabi frequencies of Fig. 3 and Fig. 4 are very well reproduced by our crystal-field model without any fitting parameter using only published values of parameters, as described above. In particular, the ratio Λ R varies from 0.58 to 0.90, whereas, in absence of orbital contribution, Λ R (m I ) = 1 for any value of m I . | cond-mat.str-el |
The orbital part of the Knight shifts is zero or negligibly small compared with their spin part since both As and La only have filled bands and/or closed shells. We estimated the total diamagnetic susceptibility derived from core electrons of each ion using the calculated values with relativistic corrections as χ dia = -6.5 ∼ -15.3 × 10 -5 emu/mol, 22,23) which is negligibly small compared with M/H. Therefore, the orbital contribution to the magnetic susceptibility is also very small in this system. Here, we would like to stress again the fact that the macroscopic magnetization of our sample can be attributed not to impurity phases but purely to the itinerant electrons of the sample, 14) since NMR can selectively detect the intrinsic signals from the sample. In general, spin parts of K and χ are related to each other as K s = A hf χ s , where A hf is the hyperfine coupling constant of the spin part at the probing nucleus. We obtained A hf values of 75 As and 139 La nuclei from the slope of K-χ plots as 75 A iso hf = 24.8, 75 A aniso hf = -7.7, 139 A iso hf = -8.64, and 139 A aniso hf = -1.41 kOe/µ B . | cond-mat.str-el |
Whether the pronounced CDW correlations observed for small λ and large ω 0 are signatures of true long-range order remains an open issue yet. To answer this question, we explore the static charge-structure factor, S c (q) = 1 N j,k e iq(j-k) f † j f j - 1 2 f † k f k - 1 2 ,(3) where 0 ≤ q < 2π. If S c (π)/N stays finite in the thermodynamic limit, CDW long-range order exists. Fig. 2 (a) demonstrates that this is the case for λ = 0.01, i.e. when the distortions of the background relax poorly. | cond-mat.str-el |
Here M + d and M - d are contributions to the spin-flip transitions due to SOI of the two lowest levels (E 0,0,±1/2 ) and higher discrete levels (E 1,0,+1/2 and E -1,1,-1/2 ). Note that the coupling to other higher levels is forbidden by the selection rule Eq. (39). The contribution of these two terms to the spin relaxation rate is shown in Fig. 8. It can be seen that these two terms interfere (constructively) which leads to a change in the amplitude and period of the oscillations. | cond-mat.mes-hall |
Fig. 9 shows the exchange splitting J and tunnel coupling t as functions of ǫ near the anticrossing point. The exchange splitting depends on both tunnel coupling and detuning, and it becomes very large near the anticrossing (resonant) point between the S(1, 1) singlet and the S(2, 0) states. In the upper (lower) panel we show the results when detuning is changed via variation of applied electric field (interdot distance). Comparison of the two panels shows that the electric field and the interdot distance have quite different effects on J and t. As the electric field increases (upper panel), the tunnel coupling increases slightly (less than 5% in the given range of detuning) because the barrier height relative to one of the dots decreases slightly. In addition, the overlap between neighboring orbitals does not depend on the electric field, so that the Coulomb terms do not change at all, and t is only affected by slightly different barrier height. | cond-mat.mes-hall |
Even though the vertex contribution is a sizable one, two remarks are in order: (1.) It does not change drastically the structure of the theory, apart from contributing towards further screening of the interactions. (2.) The fact that perturbation theory is used with the intention of being applied at a rather strong coupling is in itself questionable. Nevertheless, perturbation theory provides a clear indication that a significant contribution to screening exists beyond the conventional one-loop RPA result. On the other hand in the weak-coupling regime, α ≪ 1, RPA is parametrically well justified as far as the static polarization properties are concerned (although the RPA is not justified for the self-energy. | cond-mat.str-el |
a headless vector) or, in terms of complex field Q( r, t) = Q 11 ( r, t) + iQ 12 ( r, t) = ψ † ( r, t) (∂ x + i∂ y ) 2 k 2 F ψ( r, t). (5.2) Same as the nematic order parameter, Q is also invariant under rotations by π. The coupling between Q and N is -g N d rdt(Q † N + h.c.). (5.3) Here g N is a coupling constant. Again, the chiral symmetry of the system requires that the effective action depends only on the real part of Q † N , and that there is no dependence on the imaginary part, since it is a pseudo scalar. The tensor form of this coupling is shown in Appendix A. | cond-mat.str-el |
[19]. For the subsequent discussion it is convenient to write the current, I(t) = I (l) (t) + I (nl) (t), as a sum of linear and non-linear terms in U (t). Using Eq. ( 5) we get I (l) (t) = e 2 h T 2 ∞ q=1 R q-1 {U (t) -U (t -qτ )} , I (nl) (t) = eT 2 πτ ℑ ∞ p=1 ∞ q=1 1 p η p θ θ ⋆ r p e ipϕ(µ) R q-1 × e -iΦp(t-qτ ) -e -iΦp(t) . (6) Here R = |r| 2 and T = | t| 2 are the reflection and transmission probabilities of the QPC, k B θ ⋆ = ∆/(2π 2 ) and η(x) = x/ sinh(x). Taking the low frequency limit of Eq. | cond-mat.mes-hall |
The edge states are labeled by red bars based on the analysis we will discuss in the following paragraphs. Here we first point out that the inner edge excitations have negative ∆M and are separated by an energy gap from other states (presumably bulk states) in each angular momentum subspace. The number of these inner edge states (including the ground state) are 1, 1, 2, 3, and 5 for ∆M = 0-4, as predicted by the chiral boson edge theory. 13 They have significant overlap with a 6-hole Laughlin droplet with corresponding edge excitations embedded in a 26-electron IQH background. On the other hand, the outer edge excitations (∆M > 0) have higher excitation energies and are mixed with bulk states. In particular, for ∆M = 1, the edge state is the second lowest eigenstate in the M = 281 subspace. | cond-mat.mes-hall |
In all cases, several considerations should be taken into account such as energy spread, ion flux, vacuum environment or the eventual presence of multiple charge or mass ions. Moreover, a precise knowledge of the irradiation dose is required (direct ion current measurement with a Faraday Cup, collecting the sputtered material from the target or by measuring the etched volume by masking the target). Other relevant issues may be the unintentional temperature increase induced during the ion bombardment (target cooling is desirable) or the control over the starting material conditions (surface morphology, electrical properties, size, etc.) since small differences among experiments may lead to inconsistent results. Low-energy ion beams are difficult to handle in many cases due to spacecharge blow-up. This is normally solved by neutralizing the beam with thermal electrons from a hot filament, by placing the ion source very close to the target or decelerating the energetic ions in front of the target. | cond-mat.mtrl-sci |
The height of the layer was 3cm for glass beads and 2cm for the silica aerogel powder. The box changes its tilt angle using a mechanical system articulated on its "upper" side, while the lower side is mobile and the whole box can rotate around this side axis. The rotation is controlled through an inflated balloon. When the balloon is slowly deflated the box is rotating slowly and its position and velocity can be manipulated by stopping or regulating the quantity of air that's let out in the deflating process with a valve. The pile is softly tilted to avoid noise disturbance by the mechanical system, up to the threshold of instability where the avalanches are triggered. On the box bottom plate two Piezo-electric transducers of resonant cells (resonance frequency around 3 kHz) make the acquisition. | cond-mat.mtrl-sci |
For consistency with the previous sections, we restrict our analysis to the case of a monoclinic system, with the polarization axis, z, parallel to the heterostructure stacking direction and perpendicular to the xy plane; we shall further assume that R 1,2 are fixed, and only c (together with the ionic and electronic coordinates, {v}) is allowed to vary. Within these assumptions, the constrained-D method 13 reduces to a simpler formulation, where only the z components of the macroscopic fields D, P and E are explicitly treated. Thus, we define the internal energy functional U (D, {v}, c) = E KS ({v}, c)+ Sc 8π D-4πP ({v}, c) 2 ,(5) which depends directly on the external parameter D, and indirectly on the internal ({v}) and strain (c) variables through the Kohn-Sham total energy E KS and the macroscopic polarization P ; S = |R 1 × R 2 | is the constant cell cross-section. We then proceed to minimize the functional with respect to v and c at fixed D: U (D) = min {v},c U (D, {v}, c),(6) which yields the equilibrium state of the system as a function of the electric displacement D. D can also be expressed in terms of the reduced variable d = SD/4π, which has the dimension of a charge and can be interpreted as d = -Q free , where Q free is the free charge per surface unit cell stored at a hypothetical electrode located at z = +∞. 21 Since the surface areas of the parallel plate capacitors considered in this study are not allowed to vary, constraining D or d is completely equivalent. However, for reasons of convenience, we shall adopt d as our electrical variable in the remainder of this work. | cond-mat.mtrl-sci |
Co (0 0 0 1)[1 0 -1 0] // Al 2 O 3 (0 0 0 1)[1 1 -2 0] This growth mode holds for all films in which crystalline reflections can be observed, indicating, that the epitaxial relation dictates the grain orientation as soon as the substrate temperature is sufficiently large to result in a film with hexagonal crystal structure. The texture spread as determined from the full width at half maximum (FWHM) of the (1 0 -1 1) pole manifests a rather small tilt in out-of-plane direction of ∆Ψ = 3° and a rotational variation in-plane of ∆Φ = 6°. The epitaxial growth has to be understood from the matching symmetry between the 6 fold Al 2 O 3 surface and that of the hexagonal basal plane of hcp-Co, and from a tolerable lattice mismatch and justifies the choice of the Al 2 O 3 (0001) substrate. Whether these conditions are decisive cannot be decided from this study. Interestingly, in other works reporting hcp-Co films with perpendicular c-axis orientation the films were grown on Ru(0 0 0 1)[1 0 -1 0] // Al 2 O 3 (0 0 0 1)[1 1 -2 0] The lattice mismatch between Ru and sapphire is -1.5%. 10 The subsequent Co film now grows in a 1:1 fashion onto the Ru buffer, as can be seen in Fig. | cond-mat.mtrl-sci |
It is believed that next-generation high-performance computers can be achieved through using the spin freedom of electron in key materials and devices of current semiconductor technology [1,2]. The half-metallic ferromagnet, first discovered in NiMnSb by de Groot et al in 1983, has almost 100% spin polarization near the Fermi level [3][4][5]. This feature of half-metallic ferromagnetism makes carriers have high spin-polarization near the Fermi level and avoid some spin-related scattering processes that should exist otherwise. These are essential to practical spintronic applications [1,2,5]. Halfmetallic ferromagnetism has been found in many materials, such as Heusler alloys [3,4,6], transition-metal oxides [7][8][9], and even graphene nanoribbons under electric field [10]. The half-metallic ferromagnetic materials compatible with semiconductor technology is believed to be promising candidates for achieving more powerful computers. | cond-mat.mtrl-sci |
Hence, this mode dominates the low-energy physics and the scaling behavior. The fact that q x , q y , and ω have different scaling dimensions is typical of anisotropic systems. For instance, in the smectic (stripe) phase of the quantum Hall state, a similar scaling was found in Ref. [97]. (This scaling behavior was later on proved to be unstable [141] due to the existence of an infinite set of marginal operators in that system.) Although the problem we are discussing here and the smectic quantum Hall state share the same scaling dimensions, they are actually quite different. | cond-mat.str-el |
( 37) extends over time, space, and spin variables. The Hermitian z i matrix can be brought into the form z i = |x i1 | 2 z i1 + |x i2 | 2 z i2 x i1 y * i1 z i1 + x i2 y * i2 z i2 x * i1 y i1 z i1 + x * i2 y i2 z i2 |y i1 | 2 z i1 + |y i2 | 2 z i2 ,(39)where x i1 y i1 = 1 C i- p ix -ip iy p i -p iz ,(40) x i2 y i2 = 1 C i+ p ix -ip iy -p i -p iz ,(41) are the eigenvectors of p i , pi with p i = | p i | = |p ix | 2 + |p iy | 2 + p 2 iz ,(42) C i∓ = [2p i (p i ∓ p iz )] 1 2 ,(43)and z i1 = (1 -|d i | 2 ) -1 2 (p i0 + p i ) 2 -1 2 × 1 √ 2 e i (p i0 + p i ) + d i (p i0 -p i ) × (1 -e 2 i ) -1 2 (p i0 -p i ) 2 -1 2 ,(44) z i2 = (1 -|d i | 2 ) -1 2 (p i0 -p i ) 2 -1 2 × 1 √ 2 e i (p i0 -p i ) + d i (p i0 + p i ) × (1 -e 2 i ) -1 2 (p i0 + p i ) 2 -1 2 . (45)Then we get z i↑↓ = x i1 y * i1 (z i1 -z i2 ) ,(46) z i↓↑ = x * i1 y i1 (z i1 -z i2 ) . (47) We note that only for the half-filled band case (n ↑ + n ↓ = 1, e i = |d i |), we find that z i1 = z i2 = z i , i.e.,z i = z i τ 0(48) becomes diagonal, and the matrix elements of the original Hamiltonian are reproduced by the slave-boson transformed model. That means, Eq. ( 36) with Eqs. | cond-mat.str-el |
Therefore, we take a staggered magnetic field ∆ AF along the x axis, H AF = ∆ AF R e iQ•R (c † R,↑ c R,↓ + c † R,↓ c R,↑ ),(1.32) and consider a long-range spin Jastrow factor J J = exp 1 2 R,R ′ v R-R ′ S z R S z R ′ ,(1.33) v R-R ′ being variational parameters to be optimized by minimizing the energy. The Jastrow term is very simple to compute by employing a random walk in the configuration space |x = c † R1,σ1 . . . c † RN ,σN |0 defined by the electron positions and their spin components along the z quantization axis, because it represents only a classical weight acting on the configuration. Finally, the variational ansatz is given by |pBCS + AF = J P Sz=0 P G |BCS + AF ,(1.34) where P Sz=0 is the projector onto the S z = 0 sector and |BCS + AF is the ground state of the Hamiltonian (1.30). | cond-mat.str-el |
. 1.0t, Ohashi et al. 26 found a much larger value, T c ≈ 0.3t ≈ 140 K. The experimental data and the present ED/DMFT results suggest that the metalinsulator coexistence region is located at temperatures below those considered in Ref. 26 . For t ′ = 0.8t, the first-order phase boundaries separating the Fermi liquid from the Mott insulator in Fig. 1(a) show the same kind of reentrant behavior as measured for κ-Cl. | cond-mat.str-el |
Mesoscopic investigation by Scanning Hall Probe Microscopy (SHPM) has shown coexisting antiferromagnetic (AFM) and ferromagnetic (FM) phase around critical field in doped CeF e 2 and Gd 5 Ge 4 9,10 . For the T chosen in these studies, it did not matter whether the measurement temperature is reached by cooling or by warming. Here we present real space magnetic imaging study by SHPM along with magnetization and resistivity measurement of M n 1.85 Co 0.15 Sb to show that field induced transition for T (=120 K) lying between supercooling and superheating spinodal depends on the path followed to reach the measurement temperature. Doped M n 2 Sb shows first order antiferro (AFM) to ferrimagnetic (FRI) transition at low temperature 11 . Below transition temperature (T N ) AFM to FRI transition can be induced with the application of magnetic field. When T N is shifted to lower temperature, these systems show anomalous magnetic behavior 12,13 . | cond-mat.str-el |
We might as well conclude that in such a structure, in the absence of magnetic field, when a localized state is completely wrapped by a expanding state, the average electron occupation number in it will be close to zero and such a state will become vacuum. Thus, under the condition of U 1 = 2Γ and U 2 = Γ, the average occupation number of σ-spin electron in QD-2 has its maximum 1 2 [see Fig. 5(c) and Fig. 6(b)], since in such a case the Coulomb-induced levels in the two QDs are the same as each other. Lastly, we have to mention the influence of the electron interactions on the Fano interference when the electron correlation is taken into account. Well, when the electron interaction is very strong, one need extend the theoretical treatment by adding the interdot interaction and beyond the second order approximation. | cond-mat.mes-hall |
The difference between band and Mott insulators is also seen in the DMRG calculation of spectral function, 13 where we have needed a large number of local phononic states for band insulators in comparison with Mott insulators with the same EP coupling constant. Summarizing, we have studied the relaxation dynamics of photocarriers in the 1D Mott insulators with EP coupling. The EP coupling dominates the spin-charge coupling in the initial relaxation, even if the optical conductivity is not affected by the EP coupling. We discussed the difference of relaxation between Mott and band insulators combining the present results with U -dependence on polaron formation. This work was supported by the Next Generation Supercomputing Project of Nanoscience Program and Grant-in-Aid for Scientific Research from MEXT (19052003, 21740268, 22340097). A part of numerical calculations was performed in the supercomputing facilities in ISSP, University of Tokyo, YITP, Kyoto University, and IMR, Tohoku University. | cond-mat.str-el |
The two representations are therefore equivalent only if ǫ {R,R ′ } f R1,R ′ 1 . . . f Rn,R ′ n = f bos R1,R ′ 1 . . . | cond-mat.str-el |
From injection to detection, there are three electrically coupled regions of different spin transport: injection from the ferromagnet through the tunnel barrier via pinholes, spin transport through the graphene in the immediate vicinity of the injection point (drift region) and finally, diffusion/relaxation towards the detection point. In the drift region the DC bias gives rise to the electric field E. This yields a drift-diffusion type of transport as described in Ref. [10] , however in this case on the short length scale L. The carriers in the graphene drifting away from the injection point (in case of positive DC bias and p-type graphene) reduce the backflow of spins and thus facilitate further injection of spin polarized current through the pinhole(s). The upper limit of the effect is a measurement of the intrinsic spin polarization P possible to inject from the ferromagnetic electrodes (large E), when impedance mismatch is eliminated. On the other hand, an opposite electric field polarity (or carrier type) will result in a carrier drift towards the injection point and therefore, it will keep the local spin polarization high, reducing the efficiency of the spin-polarized injection. An increased negative bias enhances the impedance mismatch to a point when the drift starts to dominate the spin transport. | cond-mat.mes-hall |
7(a)] at high temperatures. Basically, ∆ < 0 produces a large enhancement of the susceptibility that is not strongly dependent on the temperature down to T ∼ 0.3J. 32 Hence, ∆ < 0 cannot provide an explanation for the sharp increase seen experimentally in the powder susceptibility. Notice also that in presence of an easy-axis exchange anisotropy, the in-plane susceptibilities are larger than the out-of-plane ones, and their anisotropy [inset in Fig. 7(b)] is nonmonotonic in temperature. III. | cond-mat.str-el |
Singlet-triplet intersystem crossing can occur either from the zero-point vibrational level of S 1 or from thermally-populated vibrational level of S 1 into an excited vibrational level of T 1 , or more probably into a higher excited triplet state T 2 , which is closer in energy to S 1 . It has been found [20,21] that the properties of the triplet states directly impact device performance. For example, the formation of triplet states may cause the loss of the device efficiency in these materials and thus can limit device performance and operational life span. Therefore, investigation of triplet excitations is crucial for a full understanding of electroluminescence behavior of conjugated organic polymers and for the improvement of new materials. Monkman and collaborators [21,22] investigated the photophysics of triplet states in a series of conjugated polymers and measured the excitation energies of the lowest singlet-and triplet-excitated states. Their measurements show that the excitation energies in general respect the well-known rule of thumb found for small molecules:E T ≈ 2E S /3,(1) where E T is the triplet excitation energy and E S is the singlet-singlet excitation energy. | cond-mat.mtrl-sci |
We study nuclear spin dynamics in a quantum dot close to the conditions of electron spin resonance. We show that at small frequency mismatch the nuclear field detunes the resonance. Remarkably, at larger frequency mismatch its effect is opposite: The nuclear system is bistable, and in one of the stable states the field accurately tunes the electron spin splitting to resonance. In this state the nuclear field fluctuations are strongly suppressed and nuclear spin relaxation is accelerated. Electrons confined in semiconductor quantum dots are being investigated intensively in recent years. Much research is inspired by the possibility to use their spin to implement qubits, i.e. | cond-mat.mes-hall |
In Fig. 17 we display the data obtained on Si by our group [72] and by the group of Frost [199] in terms of the normalized ordered domain size (i.e, the ratio of the lateral correlation length to the pattern wavelength) versus ion dose. For the rotating Si target the order seems to increase with ion dose irrespective of the ion species, although saturation is observed for 1 keV Kr + ions but not for 0.5 keV Ar + . In principle, it appears that the order of the pattern is larger for the rotating substrate configuration than for the fixed configuration when Ar + ions are employed. However, it should be noted that for 1 keV Ar + ions the order for the rotating configuration dropped to less than half that obtained for 0.5 keV Ar + ions. Fig. | cond-mat.mtrl-sci |
First, we rewrite Eq. ( 30) as G c σ (t, t ′ ) = G 0,σ (t, t ′ ) + C ds 1 C ds 2 G 0,σ (t, s 1 ) i n i,j=1 δ C (s 1 , t i )[(e Γσ -1)N σ ] i,j δ C (s 2 , t j ) mc G 0,σ (s 2 , t ′ ),(33) where the variables s 1 and s 2 run over the contour C, and • mc denotes the Monte Carlo averaging. It is therefore sufficient to accumulate the impurity system T -matrix X σ (s 1 , s 2 ) = i n i,j=1 δ C (s 1 , t i )[(e Γσ -1)N σ ] i,j δ C (s 2 , t j ) mc ,(34) as mentioned in Ref. 42. While the measurement of X on some fine grid introduces discretization errors, these can be made negligibly small at essentially no computational cost. Furthermore, comparison of Eq. | cond-mat.str-el |
This is just the relation q = (θ/2π)g derived by Witten 14 , for the electric and magnetic charges of a dyon inside the θ vacuum, with θ/2π = ±P 3 here. The physical origin of the image magnetic monopole is understood by rewriting part of the Maxwell's equation as ∇ × B = 2αP 3 δ(z)n × E. (4) z 1 1 , ! 2 2 , ! x y x (0,0, ) d (0,0, ) d " 1 1 ( , ) q g 2 2 ( , ) q g q TI FIG. 1: Illustration of the image electric charge and the image monopole of a point-like electric charge. The lower half space is occupied by a topological insulator (TI) with dielectric constant ǫ2 and magnetic permeability µ2. | cond-mat.mes-hall |
The angles between spin pairs are given by θ 14 = θ 23 = 2L 3 . From Eq. (1.25), this angle is related to the pitch by 2L 3 = πδ. The spiral magnetic state has the structure L A 1 = cos (πδ/2) (cos (2πδy), 0, sin (2πδy)) e 2πiz , L A 2 = 0,(1.33) L A 3 = sin (πδ/2) (-sin (2πδy), 0, cos (2πδy)) e 2πiz , producing Bragg scattering at wavevector q = 2π(0, ±δ, 1), while the ordered magnetic moments lie in the ac-plane [Fig. 1.6(b)]. All of this is consistent with the experimental data of Ref. | cond-mat.str-el |
(iv) The random distribution of the O α ions. It is usually accepted that the exchange couplings are mainly sensitive to the geometry of the coppers first coordination shells. In the CLBLCO family this geometry is controlled by the YBCO Ba sites occupation [12]. In the present work we thus focused on the effect of this degree of freedom, and averaged the electrostatic effects induced by the O α positional disorder as well as those induced by the Ca / La Y chemical disorder. We computed the effective exchange J between nearest neighbor (NN) copper atoms as well as the parameters of a Hubbard model (that is t, U and δε) as a function of the local ion distribution on the YBCO Ba site. Following the above specifications for building the embedded fragment, the quantum part was reduced to a Cu 2 O 9 fragment surrounded by 6 Cu 2+ and 10 O 2-TIPs in the CuO 2 plane, 2 Cu I TIPs on top of the apical oxygens, 6 (Ca/La) (3-x)+ averaged TIPs on the YBCO Y sites, and 6 either Ba 2+ or La 3+ on the YBCO Ba sites, as pictured in figure 1. | cond-mat.str-el |
(9) Since for each spin species, j s = σ s F s , where F s is the effective force, we finally get for the latter F ↑,↓ = ±F 16 , where F i = 2 (m × ∂ t m) • [(1 + βm×) ∇ i m] = 2 [m • (∂ t m × ∇ i m) + β (∂ t m • ∇ i m)] ,(10) after inverting the magnetic equation of motion (7) in order to express the effective field H eff in terms of the magnetization dynamics m(r, t). (Note that since the currents themselves are now generated by the magnetization dynamics, we can neglect their backaction on the magnetic response, when inverting the equation of motion to express H eff in terms of m, since it would give rise to higher-order terms.) Equation ( 10) is a key result of this paper. It is also easy to show that taking into account Gilbert damping α has no consequences for the final result (10) [after rewriting Eq. ( 7) in the Landau-Lifshitz form, in order to eliminate the ∂ t term on the right-hand side and thus make the equation suitable for the Onsager theorem]. This is not surprising, since the physics of the Gilbert damping α does not have to be related to the magnetization-particle-density coupling that determines the force (10). | cond-mat.mes-hall |
We did so by demanding K 4 = 0 (or | ∆| = 1) in S f 4 , which can be implemented by simply inserting δ(K 4 ). Here, the analogue of that additional constraint is K 2 = 0. This new condition can also be written| K + q| = K F (56) Thus, besides staying within their respective cutoffs, the choices available to K and q , when K 2 = 0, are restricted in such a way that their sum vector must sit precisely on the Fermi surface. Once K is chosen, q is obligated to connect K + q to the Fermi surface thus limiting its permissible magnitudes and angles quite severely. This is depicted in figure 2. must stay within the annulus FIG. 2: K must stay within the annulus while q must stay inside the little circle of radius Λ. | cond-mat.str-el |
It is shown that systems described by Harper's equation exhibit a Dirac point at the center of the spectrum whenever the field parameter is a fraction of even denominator. The Dirac point is formed by the touching of two subbands, and the physics around such point is characterized by the relative field only, as if the latter were null at the reference value. Such behavior is consistent with the nesting property conjectured by Hofstadter, and its experimental verification would give support to such hypothesis as well as the Peierls-Onsager ansatz used to arrive at Harper's equation when crystalline electrons move in a uniform external magnetic field. The relative simplicity with which graphene -carbon single layer sheets -may be made and handled in the laboratory has drawn much attention to the physics of massless Dirac particles. 1 In this material electrons moving in two dimensions (2D) near the Fermi level are subject to an effective energy dispersion law proportional to momentum rather than the usual momentum squared. The dynamics is similar to that of photons and phonons, except that in graphene the particles are charged fermions. | cond-mat.mes-hall |
Conductance takes place approximately at the charge degeneracy points, involving, say, electrons 0 and 1, when the energy ǫ 0 to add the first electron to the empty dot is ǫ 0 ≈ 0 (or when ǫ 0 + U ≈ 0, to add a second electron with opposite spin). Away from these points only cotunneling, processes involving two or more simultaneous tunnelling events, occur and the conductance is expected to be suppressed. However, when N = 1, the Kondo effect allows for a substantial current flow if T T K , reaching a maximum conductance of G max = 2e 2 /h, the unitary limit, at T = 0. Strictly speaking, the Kondo regime is limited to the the parameter region -U + Γ ǫ 0 -Γ [7,10]. If Γ/∆ǫ 0.5, one single orbital state is involved andT K (ǫ 0 ) = √ ΓU 2• exp πǫ0(ǫ0+U) ΓU: with this definitionG(T K ) = G max /2 , with G max = 2e 2 /h if the barriers are symmetric [7]. In the following we consider single level transport. | cond-mat.mes-hall |
(For detailed comparisons, it should be kept in mind that the misfit of -0.5% reported in Ref. 3 is only about half of that in Ref. 4.) However, our theoretical curves and the experimental data of Grigoriev et al. both show a much slower saturation of the strain response with field than was found by the Landau-Devonshire approach, where d 33 was predicted to fall by about 15% already at 50 MV/m for PZT 20/80 with epitaxy constraint. Our d 33 falls by only about 2-3% over the same range. | cond-mat.mtrl-sci |
The TM data for L = L 0 provides us also with the values of parameters K. The solution of the GDMPK equation is compared with the TM result for length L > L 0 . In the TM method, the length is defined as a number of lattice sites along the propagation direction. For a given L, we find the length s in the GDMPK equation from the condition that the mean conductance (or the mean of logarithm of the conductance in strongly localized regime) found by the two methods coincides. We find in this way also the mean free path (in units of the lattice period of the Anderson model): ℓ = (L -L 0 )/s. Our results are summarized in Figs. 12345. | cond-mat.mes-hall |
We address this discrepancy by studying an epitaxial Ni 81 Fe 19 (111)/CoO(111) exchange biased bilayer by polarized neutron and x-ray scattering and reflectivity. We show that the exchange bias for an epitaxial Ni 81 Fe 19 /CoO is several orders of magnitude less that expected due to the particular domain state of the AF layer. The available coupling energy is transformed in coercivity, mediated by the magnetically disordered interface. The blocking temperature of the exchange bias appears as the blocking of the AF domains, as revealed by neutron scattering. The temperature behavior of the frozen-in and rotatable AF spins are compared to the EB field and average domain sizes. The paper is organized as follows: in Sec. | cond-mat.mtrl-sci |
( 19), ( 20) and ( 27), the spin conductivity tensor and the spin current can be obtained, in principle, through the measure of the charge conductivity within the frequency domain alone. We believe that this may provide an electrical method to detect heavy holes spin currents in presence of k-cubic Rashba and Dresselhaus SOI. V. NUMERICAL RESULTS AND DISCUSSION We start our discussion of the numerical results by considering the isotropic case first, i.e. when only one type of SOI, Rashba (β = 0) or Dresselhaus (α = 0), is present. We considered a 2DHG formed in a GaAs-AlGaAs quantum well with a heavy hole effective mass of m * = 0.51 m o and a moderated sheet hole density of n h = 3×10 11 cm -2 ; here m o is the free electron mass. For this system, the Rashba parameter has been calculated 46 to be α = 7.48 × 10 -23 eVcm 3 , which gives rise to a HH spin-splitting at the Fermi energy of ∆ R = 2αk 3 F ≃ 0.38 meV. | cond-mat.mes-hall |
(9 Spin part We consider the magnetization, or local spins, of fixed magnitude S and varying direction n, and parametrize it by the polar coordinates (θ, φ) as (Fig. 2.1) S = S(sin θ cos φ, sin θ sin φ, cos θ) ≡ Sn. ()10 Deferring the effect of damping (friction) to the next subsection, the spin part of the Lagrangian is given by L S = S d 3 x a 3 φ(cos θ -1) -H S ,(11) where H S is the Hamiltonian of local spin, which we will specify later. The first term is known as the 'kinetic potential', and describes the spin dynamics governed by a torque equation. It has the same form as the spin Berry phase in quantum mechanics, but here we treat localized spins as classical objects. In fact, the equation of motion is derived from L S asṠ = γB s × S,(12) where γB s ≡ δH S δS is the effective magnetic field acting on localized spin (in the absence of conduction electrons). | cond-mat.mes-hall |
. , wĪ N 2 ). (20) Each subdeterminant is of the form of a Vandermonde determinant, and therefore det M = (-1) N1(N1+1)/2 × I∈P m>n (T Im -T In ) m>n T Īm -T Īn . (21) Secondly, the determinant det M can be evaluated by adding in Eq. ( 16) the first N 2 rows to the last N 2 rows. This yields det M = det -v 1 v 2 -v 3 v 4 -v 5 v 6 . | cond-mat.mes-hall |
In the latticeindependent diagrammatic representation, we denote such diagrams as These diagrams are very important for two reasons: • This class of diagrams includes the most compact terms (i.e. involving short loops) of the effective Hamiltonian that are obtained by fusing identically shaped diagrams exactly at the same place (e.g. the first term in the r.h.s of Eq. (24a)). These processes are also the dominant terms in the α-expansion of Ĥeff . Indeed, the leading order for a length-L connected process in Ĥeff = Ô-1/2 Ĥ Ô-1/2 is α L-2 . | cond-mat.str-el |
The astounding correlations between the temperature evolution of the AF domain size, frozen-in spins, and value of exchange bias are shown in Fig. 5. The temperature dependence of the H EB and of frozen-in spins is correlated with the average AF domain size and orientation. Characteristic features of three different models can be inferred from these data: the origin of the blocking temperature can be described by the M&B model, the formation of AF domains are intrinsic to the DS and Malozemoff models, and the linear dependence between the AF frozen spins and the AF domain sizes are characteristic to the Malozemoff model, demonstrating their limitations. These features, including the noncollinearity between the AF and FM spins, can all be accounted for by the Spin Glass (SG) model [9,40]. VI. | cond-mat.mtrl-sci |
We adopted a larger basis set on the central oxygen with [4s + 4p + 3d] to properly represent hole localization effects. The examined nanocluster has a ionic character, with the central oxygen atom in the Si 2 0 unit found in the Si +1.49 O -1.20 charge configuration. We found that the Si 2 O 7 H 6 nanocluster correctly reproduces one electron properties, localization effects and the charge population of solid SiO 2 . In table (1) we have reported the energies, referred to the vacuum level, and the relative decay rates of the most intense fifteen transitions, all with double holes localized on the central oxygen atom, obtained after switching on the interchannel coupling. IV. EXPERIMENTAL DATA HANDLING Since first-principles methods necessarily involve approximations, the most important of which, in scatter-ing calculations, is the approximate form of the electron wavefunction in the continuum, it is important to validate the calculations against experimental data. | cond-mat.mtrl-sci |
For a fully chaotic (or ergodic) system, electrons will access all the states in the system specially when the coupling to the leads is very weak and the electron spends sufficient time in the system. However, for a non-chaotic (or non-ergodic) system, all the states will not be accessed. Only part of the states will be accessed that will depend on the position and the details of the leads and the system, that will constitute the partial density of states (PDOS). So the contribution of these electrons to thermodynamic observables like quantum capacitance of the system, or heat capacity of the system, as well as the linear response non-equilibrium effects will be determined by this PDOS. This PDOS cannot be determined from the Hamiltonian of the isolated system as the PDOS depends on initial conditions (that is through which lead the electrons enter the system and through which lead they leave as well as the characteristics of the leads and the system). However, the scattering matrix depends on these factors. | cond-mat.mes-hall |
Measurements on VA-SWNT assisted by calculations on graphene-based systems can hence discern the contributions of the building blocks and their interaction, and show that the study of a prototype system of this kind can be used to obtain insight into the collective electronic excitations of related materials. Acknowledgements: This work was supported by the DFG PI 440 3/4, the EU's 6th Framework Programme through the NANOQUANTA Network of Excellence (NMP4-CT-2004-500198) and by the ANR (project NT0S-3 43900). Computer time was provided by IDRIS (project 544). C. K. acknowledges the IMPRS for Dynamical Processes in Atoms, Molecules and Solids. R. H. thanks the Dr. Carl Duisberg-Stiftung and C'Nano IdF (IF07-800/R). We thank S. Leger, R. Hübel, and R. Schönfelder for technical assistance. | cond-mat.str-el |
The generalized Kadanoff-Baym ansatz makes the calculational procedure in principle selfconsistent because it connects the kinetic Green's functions in the correlation contributions with the retarded (equilibrium) Green's functions. It also ensures that the calculation deals exclusively with two-time quantities, which in equilibrium depend only on the time difference, so that after Fourier transformation both the one-particle and two-particle correlation APPENDIX A: THE f -SUM RULE FOR FINITE SYSTEMS In this Appendix we show how the f -sum rule for the density-density correlation function or, equivalently, for the inverse dielectric function for finite systems follows from particlenumber conservation. To this end, we define the quantity C( r 1 , r 2 ; t 1 ) where the angular brackets indicate an equilibrium averaging, and ρ(1) = ψ † (1)ψ(1) is the particle density operator expressed by the field creation and destruction operators. We then evaluate Eq. (A1) using the particle-number conservation condition, and finally relate C to the Fourier transform of the density-correlation function. To evaluate Eq. | cond-mat.str-el |
[14]. In this scheme, only the difference between the spherically averaged screened Coulomb energy U and the exchange energy J is important for the total LDA (GGA) energy functional. Thus, in the following we label them as one single effective parameter U for brevity. In our calculation, we use J =0.51 and 0.75 eV for the exchange energies of U and Pu, respectively, and the effective Hubbard U are 4.0 and 3 eV, which are close to the values used in other previous work [4,5]. As for the optical spectra calculations, we adopt two different methods to determine the macroscopic static dielectric constants using different approximations [15]. One method is using a summation over conduction band states and the other is using the linear response theory (density functional theory). | cond-mat.str-el |
The ground state degeneracy of the featureless Mott insulator state on a torus is thus predicted to be equal to twice the ground state degeneracy of the corresponding Laughlin state, i.e. 4 in the case of filling factor 1/2. By the same logic, at odd-denominator filling factors, such as ν = 1/3, the ground states of (21) should be compressible but non-superfluid liquids ("Bose metals"), 28 corresponding to composite fermion Fermi liquid ground states of 2D bosons in magnetic field. 29 All these predictions are testable by either quantum Monte-Carlo simulations, since Eq. ( 21) does not have a sign problem, or by exact diagonalization of (21). While we have demonstrated the FQHL to featureless Mott insulator connection for the case of interacting bosons, we believe that our conclusions also hold, with possible minor modifications, in the case of interacting fermions as well, since the physics of the FQHE and of Mott insulators does not depend significantly on the statistics of the particles. | cond-mat.str-el |
3: Black(upward triangles), red(downward triangles) and green(diamonds) curves show time-averaged conductance results as a function of source-drain voltage for Ω=0.07Γ, Ω=0.14Γ and Ω=0.35Γ respectively at T=0.002Γ. FIG. 4: Black(upward triangles), red(downward triangles), blue(diamonds) and green(squares) curves show time-averaged conductance results as a function of source-drain voltage for T=0.002Γ, T=0.005Γ, T=0.02Γ and T=0.1Γ respectively at Ω=0.14Γ. conductance at zero bias gets suppressed alongside with the enhancement at V=Ω as the Kondo effect is gradually quenched. At the highest temperature the enhancement becomes indiscernible. The Kondo resonance even at zero bias is not developed sufficiently at this temperature hence the reduction of the conductance as a function of the bias is not as dramatic as lower temperatures. | cond-mat.mes-hall |
Integrability of the NLσM allows for the calculation of exact form factors and hence predictions for the detailed shape of the 2-particle contribution to S(k, ω) near k ≈ 0 and the 3-particle contribution near k ≈ π. 11,12 Since the field theory is based on S → ∞ and is only valid for k very close to 0 and π it is unclear how well any of these predictions should describe the S=1 case. Various comparisons of NLσM predictions with numerical results have been performed before, but these have necessarily focused primarily on equal time correlators and thermodynamic quantities. Experiments on quasi-1dimensional antiferromagnets have clearly confirmed the Haldane gap but the 2-particle nature of the small k excitations and the existence of a gap at k = π from the single particle excitation at ∆ to the bottom of the 3particle continuum at 3∆ have not been confirmed and have led to some questioning of the validity of these field theory predictions. One of the purposes of this paper is to compare our tDMRG calculations of S(k, ω) with the NLσM predictions. In addition, we will determine k c and examine the behavior near this wave-vector. | cond-mat.str-el |
These states are denoted as |N, S, S z , where N = i n i is the total electron number in the grain. FIG. 10: The filling of one-particle levels in the majority and minority bands of the grain with (a) N0 electrons, and (b) N0 + δN electrons, where δN = δNa + δNi, and δNa and δNi are the excess electrons added to the majority and minority band, respectively. For the ground branch with N0 electrons in the grain, the magnitude of the angular momentum S = S0. E S 0 F a , and E S 0 F i are the highest occupied levels of the majority and minority bands for the ground branch for N = N0 and S = S0. δa and δi is the level spacings near the E S 0 F a and E S 0 F i , respectively. | cond-mat.mes-hall |
We consider active rotation, that is rotation that move the points and not the reference frame. For example the rotation through an angle ψ about the z-axis is represented by R z (ψ) = cos ψ -sin ψ 0 sin ψ cos ψ 0 0 0 1 . After an active rotation R, the coordinates (r 1 , r 2 , r 3 ) of the vector r are transformed into the coordinates r ′ i = j R ij r j of r ′ = Rr. In a passive rotation, the reference frame is rotated: the basis vectors e i are transformed into the basis vectors e ′ i = j R ij e j . Thus, the coordinates of a point r are transformed by the inverse matrix: r ′ = R -1 r. To describe the transformation of the properties of a crystal under rotation, we consider the case of its charge density ρ(r). After a rotation changing r into r ′ = Rr, the charge density ρ is transformed into a "rotated" charge density ρ ′ of the rotated crystal. | cond-mat.str-el |
1a. To reconstruct the 3D glass structure, we embedded the 2D spherical diffraction pattern into a 3D array of 64×64×64 voxels, corresponding to O d = 2.7. Since the electron density of the glass structure is real, we also embedded the centro-symmetrical diffraction pattern into the same array. All other data points in the 3D array were set as unknowns. The 3D image reconstruction was carried out using an iterative phase retrieval algorithm (methods summary and supplementary discussion). Fig. | cond-mat.mtrl-sci |