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As found in the case of FM DOS, the gross features of the density of states remain unchanged with the La doping apart from the upward shift of the Fermi energy. Reaching LFMO, the Fermi level lands up in the dip of the three peak structured DOS, justifying the stability of the antiferromagnetic phase, as shown in the schematic diagram of Fig. 5. The antiferromagnetic state becomes energetically favorable, when the lling is such that it starts populating the Mo states in the majority spin channel of the FM DOS, which is highly localized due to the strong preference of the Mo-Fe hopping in one spin channel and not in another. The antiferromagnetic conguration of Fe spins, on the other hand, allows both Mo down spin as well as up spin electron to hop, albeit in dierent sublattices, thereby stabilizing the AFM phase through kinetic energy gain. In order to check the inuence of the missing correlation eect in GGA, we have also carried out GGA+U calculations with a typical U value 43 of 4 eV and J value of 1 eV, applied at the Fe site. | cond-mat.mtrl-sci |
1, a quantum particle may take a short cut from one well to the other without climbing the barrier. In this paper we will study a new class of MQT -the MQT in Z 2 topological order. At the beginning we give a brief introduction to Z 2 topological order. Topological order is a new type of quantum orders beyond Landau's symmetry breaking paradigm 2,3,4,5,6 , of which there are four universal properties : 1) All excitations have mass gap; 2) The quantum degeneracy of the ground states depends on the genius of the manifold of the background; 3) There FIG. 1: The scheme of a typical macroscopic quantum tunneling process. are (closed) string net condensations; 4) Quasi-particles have exotic statistics. | cond-mat.str-el |
calculated the FWHM of the reduced Raman spectrum and not its half width at high energy, ( 2 / Γ ). Second, they carried out their simulations with a-Si models whose bond angle distributions were not Gaussian-like and decided to quantify their dispersion through the width of the Gaussian distribution, G θ Δ , that best fitted the actual distribution. In other words, the curve labelled "Maley" in Fig. 1 Δ respectively, the corrected relationship of Maley is very close to that of Beeman-Tsu. This coincidence is very reassuring because, although the microscopic models used by Maley et al. [14] were similar to those of Beeman et al. | cond-mat.mtrl-sci |
For a short and wide strip there is a large number W/L ≫ 1 of evanescent modes with transmission probability of order unity. In a remarkable coincidence,7 the transmission probabilities of the evanescent modes are the same as those of diffusive modes in a disordered piece of metal with the same conductance (Tworzyd lo et al., 2006). We will return to this "pseudo-diffusive" dynamics in Sec. III.D, when we describe how supercurrent flows through undoped ballistic graphene in the same way as it does through a disordered metal. In preparation of that discussion, we examine here in a bit more detail the transmission of evanescent modes through undoped graphene (Katsnelson, 2006b;Tworzyd lo et al., 2006). Because the wave length at the Dirac point is infinitely long, the detailed shape of the electrostatic potential profile at the interface between the metal contacts and the graphene sheet is not very important. | cond-mat.mes-hall |
(46,47). We mention that within first order in SOI the decoherence time T 2 induced by phonons satisfies T 2 = 2T 1 since, as mentioned before, the fluctuating magnetic field induced by phonons δB is perpendicular to the applied one B. In Fig. 2 we plot the relaxation time as a function of the ratio ω eff Z R/c l , for R = 50 nm and c l = 4 • 10 3 m/s. We see that the relaxation rate exhibits peaks as a function of the effective Zeeman splitting E eff Z . This is due to the finite size in the transverse direction which gives rise to phonon branches. | cond-mat.mes-hall |
( 6), show the following: At high temperatures I (nl) (t) vanishes and we find (T > t -t 0 > 0): I(t) = e 2 δU T R N (t) /h, θ ≫ θ ⋆ ,(7)where the integerN (t) = [[(t -t 0 )/τ ]]. The current decreases with time in a step-like manner being constant over a time interval τ . At t -t 0 ≫ τ we can write I(t) ∼ e -(t-t0)/τD with decay time τ D = h/(∆ ln(1/R)). This agrees well with Ref. [2] since τ D ≈ τ unless T ∼ 1. Eq. | cond-mat.mes-hall |
We have investigated the magnetic and dielectric properties of Ni 3 V 2 O 8 in great detail 8-10 . Ni 3 V 2 O 8 has a layered structure, with the spin-1 Ni 2+ ions sitting at the vertices of distorted Kagome planes stacked along the (010) axis 8 . Ni 3 V 2 O 8 has a rich phase diagram: at zero magnetic field, there is a transition from the disordered phase into a high temperature incommensurate phase (HTI) at 9.3 K, followed by a transition into a low temperature incommensurate phase (LTI) at 6.3 K, and there is a canted antiferromagnetic (CAF) phase with a small ferromagnetic moment along (001) below 3.9 K. The phase diagram is sensitive to an external magnetic field; applying an external magnetic along (100) promotes the LTI phase, while a field along (001) favors the CAF phase 9 . Because Ni 3 V 2 O 8 is insulating with a small ferromagnetic moment at low temperatures, we might expect it's dielectric constant to show an anomaly at the magnetic transition temperature. This assumption is confirmed in the data presented in Fig. (2). | cond-mat.mtrl-sci |
The energy added to/removed from the system ( R V C (q)dq) is the area between the curve and the q axis. The areas of shaded regions U 1 and U 2 give the amount of added/removed energy in each half-period. The signs of U 1 and U 2 are determined by the direction on the loop. For the direction shown here, U 2 is positive and U 1 is negative. and a charge-controlled memcapacitor when Eqs. ( 12) and ( 13) can be written as q(t) = C t t0 V C (τ ) dτ V C (t) (14) q q U 2 V C U 1 U 1 Fig. | cond-mat.mes-hall |
Consider the low-energy excitations with momenta in the vicinity of the Fermi level near the K point in the Brillouin zone of graphene which obey a Dirac-type equation [10] 0 ) , , ( ) , ,( ) ( 0 0 = ⎟ ⎟ ⎟ ⎠ ⎞ ⎜ ⎜ ⎜ ⎝ ⎛ ⎟ ⎟ ⎟ ⎠ ⎞ ⎜ ⎜ ⎜ ⎝ ⎛ Δ + + ∂ - ∂ + ∂ ∂ - ∂ Δ - + ∂ t y x t y x y qE i i i y qE i t y x y x t χ φ (1) where (Fermi velocity) and are set to unity and a static uniform electric field -in the y-direction is present. We are working in the space-dependent Coulomb gauge so the vector potential is . The matrices operate on the one-valley, triangular sublattice (pseudospin) space of the graphene honeycomb structure corresponding to the A and B atoms. Although the fermions are massless we have introduced a term for later convenience. Each Fourier mode of the field can be expanded as , so Eq. ( 1) simplifies into two Schrodinger-type equations for a unit mass in an inverted harmonic oscillator potential F v ) t = h ) 0 ,t y 0 E , 0 , ( 0 y E A = μ ) ( y a , (x χ Δ , , ( ) ( e y x t x k i x ω φ - ) ( ) ) ( y b e t x k i x ω - = 0 ) ( ) ( 2 4 1 2 2 2 = ⎟ ⎟ ⎠ ⎞ ⎜ ⎜ ⎝ ⎛ ⎟ ⎟ ⎠ ⎞ ⎜ ⎜ ⎝ ⎛ ± - + ⊥ ξ ξ ξ ξ b a i a d d k (2) in which ) ( / 2 0 0 y qE qE + = ω ξ , . | cond-mat.mes-hall |
1) implies that the lowlying excitations arise from nonmagnetic origin. Next, we analyze the low-lying excitations using the temperature dependence of 1/T 1 T at the La site. We consider that 1/T 1 T at the La site is governed not only by the Korringa mechanism but also by local fluctuations of the EFG due to the rattling motion of the La ions. In fact, ultrasonic dispersion is found around 50 K in LaOs 4 Sb 12 , 21 which is considered to originate from the rattling motion of the La atoms as in PrOs 4 Sb 12 . 7 We assume that the fluctuations of the EFG disappear at low temperatures and that the Korringa mechanism dominates the relaxation at the La site at low temperatures, because 1/T 1 T at the La sites becomes constant below 10 K down to 0.6 K as 1/T 1 T of 121 Sb does as shown in the inset of Fig. 4. | cond-mat.str-el |
However, we also need higher order terms in Eq. ( 16), which will arise from including the fluctuations of the gapped fields Q and λ. Rather than computing these from the microscopic Lagrangian, it is more efficient to deduce their structure from symmetry considerations. The representation in Eq. ( 14), and the connection of the U , V , W to the lattice degrees of freedom, allow us to deduce the following symmetry transformations of the X, Y , Z: • Under a global spin rotation by the SU(2) matrix g σσ , we have Z σ → g σσ Z σ , and similarly for Y , and Z. When DM interactions are included, the global symmetry is reduced to U(1) rotations about the z axis, under which Z ↑ → e iθ Z ↑ , Z ↓ → e -iθ Z ↓ Y ↑ → e iθ Y ↑ , Y ↓ → e -iθ Y ↓ X ↑ → e iθ X ↑ , X ↓ → e -iθ X ↓ . | cond-mat.str-el |
We note that a decrease of Ca-contents in the bulk of (La 1-x Ca x )MnO 3 film and/or eventually at the film/substrate interface leads to an expansion of the unit cell due to the larger ionic radius of La 3+ (≈ 1.16 Å) compared to that of Ca 2+ (≈ 1.12 Å). Therefore, we could expect that the elastic energy required to accommodate lattice mismatch (tensile in the present case) can be reduced by compositional changes. In the present case, enrichment of the La/Ca ratio deep into the film is foreseen. This process would lead to Ca 2+ enrichment at the film surface in agreement with our XPS measurements. A similar process would also be operative in La 1- x Sr x MnO 3 thin films on STO substrates, which are also under tensile strain. In fact, as reported by Maurice et al. | cond-mat.mtrl-sci |
Here we have shown that it also influences the value of spin-orbit splitting in CNTs. We consider now chiral NTs (n, m), with n = m = 0 that do not have an inversion center. As mentioned above, curvature effects induce a shift of the Fermi wave vector k F , opening a small gap at the Fermi energy in the primary metallic chiral CNTs, (m-n) = 3q. But, in contrast to the results shown for achiral zigzag NTs, in both metal and semiconductor chiral NTs the SO interaction lifts all degeneracies. The calculated energy splittings for these tubes follow the same behavior as those obtained for achiral zigzag CNTs: for tubes with ν = 0 (nm = 3q) and ν = -1 (n-m = 3q-1) the energy splitting is larger for the highest VB, while for tubes with ν = +1 (nm = 3q + 1) the splitting is larger for the lowest CB. As an example for chiral tubes, the band structures calculated including the SOI for three particular tubes, (6,4), (9,3), and (8,4) belonging to each of the three families, are shown in Fig. | cond-mat.mes-hall |
Hence, we see that (16) saturates the bound in (17) when J ⊥ /2 > J ; (15) is the ground state for J = J X with J ⊥ /2 > J . 30,47 We now consider the zigzag ladder. It can be decomposed into two triangles per rung: H Z = l H (2) l+ + H (1) l-;(18)whereH (2) l+ = 1 2 J ⊥ S(1) l • S (2) l + J S (2) l • S (2) l+1 + 1 2 J 2 S (1) l • S (2) l+1 ,(19)H (1) l-= 1 2 J ⊥ S(1) l • S (2) l + J S (1) l • S (1) l-1 + 1 2 J 2 S (2) l • S(1)l-1 . When J = J 2 /2, the terms in ( 19) can be written in the form of (11); the energy of the state in (15) saturates the lower bound in (17) for J ⊥ /2 > J . Hence, ( 15) is the exact ground state for J = J 2 /2 with J ⊥ /2 > J . 46 Note that another exact ground state of the zigzag ladder follows by symmetry. | cond-mat.str-el |
At first the ZB phenomena in crystalline solids was predicted with the use of LCAO method in Refs. [3-5]. Later an oscillatory motion of electron wave packets has been considered in wide class of 3D solids and nanostructures, including narrow gap semiconductors 6 , carbon nanotubes 7 , 2D electron gas with Rashba spin-orbit coupling 8,9 , single and bilayer graphene 10,11 , and also superconductors 12 . The ZB of photons near the Dirac point in 2D photonic crystals was discussed in Ref [13]. Rusin and Zavadski considered ZB in crystalline solids for nearly-free and tightly bound electrons and concluded that ZB in solids is a rule rather than an exception. 14 They determined the parameters of trembling motion and concluded that, when the bands are decoupled, elec-trons should be treated as particles of a finite size. | cond-mat.mes-hall |
5 When d = 0, since the electronhole interaction is given by V E (r) = -e 2 /ǫ √ r 2 + d 2 , the differential equation for the radial component of the ex-citonic wavefunction that cannot be analytically solved. Instead, we use the momentum-space Schrödinger equation 2 k 2 2m r ψ α (k) + k ′ V E (|k -k ′ |)ψ α (k ′ ) = E α ψ α (k) (1) where k is the two-dimensional wavevector, ψ α (k) is the momentum-space eigenfunction with eigenvalue E α , and V E (q) = -V A (q)e -qd where V A (q) = 2πe 2 /ǫq is the Fourier transform of the intralayer Coulomb interaction in two dimensions. We focus on Eq. (1) projected onto the zero angular-momentum sector and obtain the eigenenergies and eigenfunctions by discretizing the integral equation and numerically diagonalizing the resulting matrix 7 H mn = u 2 n 2 δ mn + u n ∆u 2π Ṽ (u m , u n ) = u m u n H * nm . (2) Here Ṽ (u m , u n ) is the angular-averaged dimensionless electron-hole interaction. Although the Hamiltonian (2) leads to bound and continuum states, since the excitonic internal states are only accessible at temperatures T ≥ E 0 /k B ∼ 300 K, in the following we only discuss the behavior of the ground state. | cond-mat.str-el |
Whether the transition region is abrupt or graded, the transmission is perfect when θ=0 (i.e. commonly refered to as Klein tunneling), but the transmission decreases as θ and D w increase. This angular selectivity for electron transport across the pn junction serves as a filter, allowing states with |θ|≤σ θ (where σ θ is the spread of the angular distribution) to pass through more effectively. The quantity 2σ θ can be viewed as the bandwidth of this filter and is what gives rise to the larger resistance of a pn junction as compared to a uniform graphene sheet. Several research groups have recently fabricated graphene pnp devices by using electrostatic gates to create p and n regions [12,14,15,16]. The typical setup consists of a back-gate and top-gate, which are used to control the amount of charge density in the source/drain and channel regions respectively. | cond-mat.mes-hall |
Collective spin states on single tetrahedra then replace individual spins as the fundamental basis for excitations and longer range effective interactions. The main features of the spin liquid phase follow from this scenario. Pyrochlore crystals such as Tb 2 Ti 2 O 7 belong to the cubic space group F d 3m (O h 7 , No. 227), in which both Tb and Ti ions form separate, corner-sharing tetrahedral networks. Here we are only concerned with the magnetic Tb 3+ ions. The tetrahedra appear in two different orientations (A and B), which alternate in the tetrahedral network. | cond-mat.mtrl-sci |
8 for the moments of the time t = 0.1 and t = 0.5 (in units of t 0 ) correspondingly and for a = 6. The Fig. 8(b) demonstrates clearly the splitting of the angular momentum density into two parts similar to probability density discussed in Sec.II for a >> 1. The angular momentum density changes its sign within each part that looks like as multipole structure. Sx (t) = - √ 3 2 , Sy (t) = - √ 3 2 sin 2γ 2 hk 2 0 t m , Sz (t) = 1 4 1 + 3 cos( 2γ 2 hk 2 0 t m ) . (39) IV. | cond-mat.mes-hall |
Besides the semiclassical theories, the research on diluted magnetic semiconductors also stimulated the theoretical interest in other quantitative approaches. Due to the relative simplicity of several important models, such as the Rashba coupled 2D electron system, a number of publications appeared recently with rigorous quantum mechanical calculations by Kubo and Kubo-Streda formulas [6,10,13,49,53,56,57,58,59] and by a variety of the quantum Boltzmann equation and the Keldysh techniques [8,9,60,61]. Sinitsyn et al [51,53] demonstrated the 1-1 correspondence between semiclassical contributions to the AHE and the summation of relevant subseries of Feynman diagrams in the Kubo-Streda formula [62]. Similar agreement was established with the Luttinger's theory [36]. The results are summarized in Table 1. According to [53], the classification of contributions in the Kubo formula is not merely by a separation of diagrams into the disorder free part and the vertex correction but rather by the parts of the velocity matrices in chiral basis that stay inside the trace of the Kubo formula. | cond-mat.mes-hall |
It is also interesting to examine the variation of the energy barriers for migration in the initial energy curves, shown in the figure 7. The work done by the stress during the kink migration allows to further reduce the energy barrier. For instance, the latter drops to 75 meV for the largest stress considered here. IV. DFT RESULTS A. Single kink: structure We now focus on the results obtained from firstprinciples calculations, in particular the structure of the kink. | cond-mat.mtrl-sci |
This is consistent with the conclusion in a recent photoemission study revealing |ǫǫ F | 1.5 dependence of the high resolution photoemission lineshape: electron-magnon coupling may be important. [9] The magnetic moment centered at Ir and O1 sites are about 0.27 µ B and 0.04 µ B respectively. The total moment of 0.9 µ B /fu is close to the saturation moment of 1 µ B expected for S = 1/2 state of Ir 4+ . However, this is significantly high considering that no saturation was observed experimentally even at a high field of 12 Tesla. [8,16] The magnetic moment in the paramagnetic phase was found to be about 0.55 µ B (µ = 2 S(S + 1)µ B assuming spin only value), which corresponds to S = 0.07. This is about 14% of the S=1/2. | cond-mat.str-el |
11). Such a non-linear feature in the isothermal curves is also found in ref. [24,25], and may be associated with complex magnetic phenomenon in the critical region, rather than one simple, clearly defined, Landau type, 2nd order phase transition. In contrast to the Fe column compounds, the Co column compounds all appear to order antiferromagnetically with the values of T N between 4 and 7 K. Figures 12,13 and 14 present the low temperature magnetic susceptibility, specific heat and electrical resistivity data for GdCo 2 Zn 20 , GdRh 2 Zn 20 and GdIr 2 Zn 20 respectively. In addition to these data, d(χ(T )T )/dT [26] and dρ/dT [27] have been added to the susceptibility and resistivity plots respectively. GdCo 2 Zn 20 and GdRh 2 Zn 20 manifest clear λ-type anomalies in their temperature dependent specific heat, with similar features appearing in their dρ/dT and d(χ(T )T )/dT data. | cond-mat.str-el |
Experiments have shown that, during relaxation, the bond angle dispersion and the density of point defects diminish [5,8,12,24,25]. This evolution has been simulated by molecular dynamics, too [26,27]. However, the relative contribution of both effects to the relaxation energy has been the object of controversy for many years. Whereas Stolk et al. [24] assign most of the relaxation energy to bond-angle strain, Roorda et al. [25,28] assign it to defect recombination. | cond-mat.mtrl-sci |
For instance for the octonacci chain we found β ′ = 0.81 -0.85, where some of these differences might be caused by fluctuations of the width d(t) which are present in the regime of strong expansion and make fitting difficult. These scaling exponents tend towards the exponents obtained for v → 1, 20 which indicates that the fast expansion is not governed by the weak coupling, but rather a kind of resonance between the different levels of the hierarchy. Further, for the return probability we find the exponents δ ′ Au ≈ 0.71, δ ′ Ag ≈ 0.71, and δ ′ Bz ≈ 0.76, which are again relatively close to the exponents for v → 1. 20,28 However, these values of the scaling exponent δ ′ might differ from the exact result, because in one dimension we cannot rule out the influence of subdominant logarithmic contributions for the considered short time intervals in the step-like process. 28 Similar behaviors for d(t) and C(t) have been reported before for the Fibonacci chain with strong quasiperiodic oscillations. 28,30 Wilkinson and Austin found the same step-like process when studying the spreading of a wave packet for Harper's equation of an electron in a magnetic field. | cond-mat.mes-hall |
As a result, the conductivity from conduction electrons is free from vertex corrections, becoming σ c (T ) = C N c F v c2 F 2Γ c (T ) ,(35) which coincides with that of the diagrammatic study [14] showing Γ c (T ) ∼ T ln(T /E * ) in the z = 3 critical regime. The last work is to find an actual expression for the physical conductivity, referred as the Ioffe-Larkin composition rule [26] σ(T ) = σ c (T ) + σ b (T )σ f (T ) σ b (T ) + σ f (T ) ≈ σ c (T ),(36) where σ b (T ) is the holon conductivity, much smaller than fermion contributions justifying the last approximation. One can ask the role of the spinon conductivity for any physical response functions. Actually, it contributes to the physical thermal conductivity given by the corresponding Ioffe-Larkin composition rule κ(T ) T ≈ κ c (T ) T + κ f (T ) T ,(37) where κ c,f (T ) are thermal conductivity of conduction electrons and spinons, respectively, and holon contributions are also neglected. Assuming that the Wiedemann-Franz law holds for each fermion sector, proven to be correct at least in the one loop approximation [14], we find κ t (T ) T ≈ π 2 3 σ c (T ) + σ f (T ) ,(38) suggesting that the Wiedemann-Franz law should be violated due to the presence of additional entropy carriers, that is, spinons at the Kondo breakdown QCP in the low temperature limit, i.e., L(T ) ≡ κ(T ) T σ(T ) ≈ L 0 1 + ρ f v f F ρ c v c F (39) with L 0 = π 2 /3, the value of the Fermi liquid. This result would be robust beyond our approximation because this expression includes just density of states and velocity at the Fermi energy, thus expected to be governed by a conservation law. | cond-mat.str-el |
As an extension of the BCF model, the boundary conditions for (2.13) are now formulated by linear kinetics with inclusion of both atom attachment-detachment and step permeability [22,28,41]: (2.14) f ± = D A ± (ρ ± -ρ ± 0 ) ± D ± p (ρ + -ρ -) ; cf. (1.2). Here, f ± is the adatom flux normal to an edge from the upper (+) or lower (-) terrace, i.e.,(2.15) ∓ f ± := vρ ± + D T n • (∇ρ) ± , ρ ± is the terrace adatom density restricted to the step edge, D A ± is the attachmentdetachment rate coefficient and D ± p is the permeability rate coefficient. These rates can account for different up-and down-step energy barriers, e.g. the ES effect in the case of D A ± [11,32]. The reference density ρ ± 0 is given by (1.3) where ρ * is replaced by ρ ± * for up-and down-step edge asymmetry. | cond-mat.mtrl-sci |
As the total momentum of the scattered electrons is conserved in crystal [neglecting the Umklapp processes], we search for the graphene state with zero total momentum. In this state there are on average as many holes as particles in every small momentum cell. Therefore, we can describe the particles by the distribution function in the momentum space F + (p) whereas the distribution function for the holes 1 -F -(p) has to be the same. Thus, the distribution function possesses the particle-hole symmetry: 1 -F α (p) = F -α (p). In the 'equilibrium' state of the graphene the electron distribution function: f α (p) = 1 exp(α|p|/ p ) + 1(5) makes the collision integral due to the Coulomb interaction to vanish for any scale parameter in the momentum space: p . This scale also defines the effective temperature of the electrons in the 'equilibrium' state of the graphene: T * = c p . | cond-mat.mtrl-sci |
In the RS-FeN type with a density of states at the Fermi level n(E F ) = 71.596 states/Ry/Fe and a Stoner parameter I s = 0.034 Ry the Stoner criterion is fulfilled (2.434). On the contrary, for the ZB-FeN and W-FeN we have n(E F ) = 14.106 states/Ry/Fe and n(E F ) = 15.152 states/Ry/Fe respectively, the Stoner criterion (resp. 0.48 and 0.515) is not fulfilled. As a consequence, the magnetic behavior of FeN is consistent with the Stoner theory. As it was discussed in the literature, the magnetic moment of Fe atom in FeN [10,12,13] and generally in the Fe nitrides [2] is very sensitive to nearest-neighboring Fe-N distance and their d-p hybridization. Our results confirm that this latter is the most important effect. | cond-mat.mtrl-sci |
Acknowledgments This work was supported by the Swiss National Science Foundation, the STREP project SUB-TLE, and the Swiss National Center of Competence in Research, MaNEP. Appendix We present here the expression for the charge relaxation resistance in terms of a self-consistent potential landscape. For a geometrical capacitance, in random phase approximation, R q does not depend explicitly on the interaction. However, in the more general case treated here, an explicit dependence on interaction is found. The results can be expressed in terms of an effecive Wigner-Smith (density of states) matrix. Consider a cavity with M incident channels. | cond-mat.mes-hall |
We give the solutions first for MLG and then for BLG. The results are illustrated in Fig. 7. Monolayer graphene For MLG α k = s α vk, with s α = ±1. Eliminating p in Eq. (A1), we find s β vp s (k, q, x) = s α vk + s cq,(A2) where p s (k, q, x) = |p s | = k 2 + q 2 + 2kqsx, with p s = k + sq (s = ±1), x = cos φ = q • k/(kq), and s = ±1 for absorption or emission, respectively. | cond-mat.mes-hall |
With increasing flavors, more electrons y k z k x q i i E(q)=ε k ) p ( p FIG. 2: The dark grey ellipsoids show Fermi surfaces of electrons in the six degenerate conduction band valleys in Si. E(q) is the energy with momentum q measured with respect to the Γ point, ǫi(p) is energy with momentum p measured with respect to the center of the ith valley. are within ∼ k B T of the Fermi surface hence are able to be thermally excited, therefore the heat capacity of the MFEG, C = 12k 2 B T (ν/3π 2 n) 2/3 /5, increases with number of flavors. The Stoner criterion [38,39] for band ferromagnetism states that for opposite spin electrons interacting with positive exchange energy U , ferromagnetism occurs when g(E F )U ≥ 1. With increasing number of flavors the total DOS g(E F ) ∝ ν 2/3 increases so that the Stoner criterion becomes more favorable. | cond-mat.str-el |
[126]). The reason for the success of the classical theory of melting in two dimensions is that, as in all Kosterlitz-Thouless phase transitions [5,127], at finite temperatures the classical ordered state with a spontaneously broken continuous symmetry is not possible in two dimensions. Instead, there is a line (or region) of classical critical behavior with exactly marginal operators. The defect-unbinding phase transition appears as an irrelevant operator becoming marginally relevant. In the case of the quantum phase transitions in two dimensions that we are interested in, there are no such exact marginal operators available at zero temperature, and hence, no lines of fixed points available. Thus, the T = 0 ∆ CDW Wigner crystal Fermi liquid The cross point of the two dash lines is the multi-critical point ∆N = ∆CDW = 0. | cond-mat.str-el |
An analytic continuation yields tan θ → i and φ 12 → igπjL/W in Eq. ( 4), where we use the quantization (5) and define the half-integer j ≡ l + 1 2 . As pointed out by Katsnelson and Guinea [15], the zero-energy solution of the Dirac equation may be obtained via conformal transformation that links the considered geometry to a simple one, for which the wavefunction is known [22]. In particular, if the conformal transformation z(w) turns the system under consideration into a rectangle of width W and length L (Fig. 2), the transmission probability for the j-th evanescent mode may be written as T j = 1 cosh 2 [gj ln Λ{z(w)}] = 4 (Λ gj + Λ -gj ) 2 ,(6) where j = ± 1 2 , ± 3 2 , . . | cond-mat.mes-hall |
3, where the solid and dotted curves correspond to the same meaning as earlier. The increment of the resonant widths is due to the broadening of the molecular energy levels, where the contribution comes from the imaginary parts of the two self-energies in the strong molecular coupling to the electrodes. The scenario of electron transfer through the molecular bridge system is much more clearly observed by studying the current-voltage character-istics. The current is computed by the integration procedure of the transmission function T which shows the same variation, differ only in magnitude by the factor 2, like as the conductance spectrum (Figs. 2 and3). The current-voltage characteristic in the weak-coupling limit for the molecular system is given in Fig. | cond-mat.mes-hall |
At zero temperature, this computation is simplified on two counts: 3 (i) the sums over eigenstates in Eqn. (4) reduce to a single sum; and (ii) the matrix elements needed are of a single type, namely those connecting the ground with various excited states. At finite temperature however, we must deal both with the double sum in Eqn. ( 4) and the matrix elements in their full generality. To make this task tractable, we exploit the fact that the spin chain material is gapped. On a qualitative level the excitations can be divided according to the number, n, of magnons they contain. | cond-mat.str-el |
In our case the role of Ω n is played by frequencies ν 1,2 (n, h), which are determined by Eq. (15). Therefore it is reasonable to suggest that for traveling Rabi waves the local fields effects are negligible under the conditionν 1,2 (n, h) ∆ω, (57) where ∆ω is determined by Eq. ( 8). In the case of Rabi-wavepackets general line of reasoning is the same as for Eq. ( 57) and leads to the following condition of the local-field negligibility: ν 1,2 ( n , h In this section we will calculate exciton-exciton and exciton-photon multi-time correlation functions. | cond-mat.mes-hall |
This gives the classical dielectric model. The starting point is the dielectric constant of the structure, that is either the bulk one, when quantum confinement effects (QCEs) are negligible, or a size dependent Si-nc dielectric constant, derived, for example, from a Penn-like model or a semiempirical approach [15]. In the following we start from the independent particle dielectric constant derived from our tight binding method. We refer to this model as semiclassical, since QCEs are correctly taken into account, but SPEs are included using classical electrostatics. It is worth stressing that in this model SPEs enter the polarizability of a structure only through the geometrical shape and the dielectric mismatch across the surface, but they are independent of the nanocrystal size. The dielectric model has been shown working well for huge structures with a quite regular shape. | cond-mat.mtrl-sci |
In a magnetic field, electrontransfer processes redistribute carriers to minimize the free energy and a strain is induced. Given that Bi 2 Se 3 shares this many-valley form of the band structure, a similar mechanism might provide a cohesive explanation for the simultaneous change in lattice dynamics and shift of the band edge with magnetic field. Further measurements are necessary to determine if appreciable magne-tostriction occurs in Bi 2 Se 3 . Magnetoelectric coupling uncovered by our experiments could further be related to the topological nature of the material. The topological magnetoelectric effect, in which an applied magnetic (electric) field creates an electric (magnetic) field in the same direction, is predicted to occur in a nontrivial topological insulator. 9,44 The sensitivity of infrared active phonon modes to perturbations of local electric fields is wellknown. | cond-mat.str-el |
However, the energy dissipated during the fracture process is a material property and should be independent of the discretisation applied. This limitation of local constitutive models can be overcome by adjustment of the softening modulus of the softening branch of the stress-strain curve with respect to the finite element size [13,2,19]. With this approach, the dissipated energy is modelled mesh independently, as long as the inelastic strains localise in a zone of assumed size. This approach, which was used earlier for the analysis of delamination in sandwich structures [1] and cracking in cohesive materials [7], is used in the present study. The present NLFM approach differs strongly from linear elastic fracture mechanics (LEFM) approaches, which assume a large stress-free crack with all the nonlinearities of the fracture process concentrated in an infinitesimally small zone in front of the crack tip [5]. If the size of the fracture process zone is large compared to the length of the stress free crack and the size of the structure, LEFM results in a poor approximation of the fracture process. | cond-mat.mtrl-sci |
These experiments have also established that static long-range stripe charge and spin orders do not have the same critical temperature, with static charge order having a higher T c . An important caveat to our analysis is that in doped systems there is always quenched disorder, and has different degrees of short range "organization" in different high temperature superconductors. Since disorder also couples linearly to the charge order parameters it ultimately also rounds the transitions and renders the system to a glassy state (as noted in Refs. [3,53]). Such effects are evident in scanning tunneling microscopy (STM) experiments in Bi 2 Sr 2 CaCu 2 O 8+δ which revealed that the high-energy (local) behavior of the high temperature superconductors has charge order and it is glassy [53,71,72,73,74]. Finally, we note that in the recently discovered iron pnictides based family of high temperature superconductors, such as La (O 1-x F x )FeAs [75,76], a unidirectional spin-density wave has been found. | cond-mat.str-el |
The resulting integrals are highly non-trivial, and presently we do not know how to evaluate them. VII. THE MOORE-READ PFAFFIAN STATE In the previous sections we have constructed a quasilocal quasielectron operator for the Abelian hierarchical states. The concepts and ideas behind the construction are however very general, and can, in particular, be generalized to the proposed non-Abelian states. Of particular interest among these is the Moore-Read (MR) Pfaffian state which is believed to describe the incompressible state observed at ν = 5/2. This claim is based both on extensive numerical calculations 36 and the recent observation of charge e/4 quasiparticles. | cond-mat.mes-hall |
58,66 As a rule of thumb, particle-hole conjugate states contain counter-propagating edge modes and one needs to consider the effect of edge disorder. (The ν = 2/3 state can be thought of as ν = 1/3 fluid of holes in a ν = 1 background. Thus, the edge theory in the absence of disorder can be described by an "outer" ν = 1 edge and an "inner" ν = -1/3 counter propagating edge.) The same scenario will arise in particle-hole conjugate states in non-Abelian quantum Hall systems. One must analyze the relevance of disorder there as well and determine the new effective low-energy theory of the edge if disorder is relevant. 24,25,39 B. Moore-Read Pfaffian edge state Fermi statistics requires that a large class of Abelian states have odd-denominator filling fractions, such as the most prominently observed ones, which have ν = n 2pn+1 . | cond-mat.str-el |
Thus, the spin motive force e is created in the radial direction for skyrmion spin textures. Since e zu is non-zero and there is the external magnetic field and b zu , a drift of conduction electrons is induced. The drift is circular around the skyrmion because e zu is in the radial direction. Thus, the drift motion leads to the vortex current. The drift velocity is v φ = 2αBλ 2 r 2λ 2 + (eB/h) (r 2 + λ 2 ) 2 . (14) Figure 1 shows this drift velocity as a function of r/λ. | cond-mat.str-el |
( 6)- (7), two competitive mechanisms manifest themselves in the light -QD-chain coupling: the local-field induced nonlinearity and the dispersion spreading due to the tunneling. To this point we have taken no account of the decay processes inside the QDs. To take this processes into account one should allow for processes of interaction of e-h pair in QD with phonon bath. Such interaction can be described by conception of quantum trajectories 37,38,39 of e-h pair, which are the superpositions of deterministic evolutions and random jumps under the action of Lindblad operators. In this case the Hamiltonian (1) should be replaced by the efficient non-Hermitian Hamiltonian Ĥeff = Ĥ -i p X+ p Xp , where Xp is Lindblad operator for p-th QD. Suppose that the density of phonon states of the bath has the form of Lorentz line with frequency ω 0 and width λ. | cond-mat.mes-hall |
To see how we can get a Mott type law from extended states, we use the band edge term of Eq. ( 32) and Eq. ( 37) to get σ xx = e 2 k B T εc(Bz ) -∞ dερ band D 0 ε c (B z ) -ε ε c (0) e -| ε f (Bz )-ε | k B T - " εg |ε f (Bz )-ε|+ε l " ν . (38) A steepest descent evaluation of Eq. ( 38) at low T gives for ν = 1/2, an exp -T0 T 1/3 law (see Appendix A). Here, the Mott-like laws are a result of the pseudo-gap in the density of states or a depletion of low energy excitations created by long range correlations. | cond-mat.mtrl-sci |
The dc gate voltage V g changes the gate charge en g = C g V g , which controls the tunneling of Cooper pairs to and from the island. The other parameter is the magnetic flux bias Φ b through the inductor loop caused by an external magnetic field. This controls the Josephson coupling E J of the island as the flux shifts the relative phases of the two junctions and leads to complete suppresion at Φ b = Φ 0 /2 in a symmetric SCPT. Here Φ 0 = h/2e is the flux quantum. The properties of the SCPT can be calculated in detail using Mathieu functions 26 and including the asymmetry of the junctions. Instead of the general analysis we present here a simplified treatment that is still sufficient to understand the main properties of the circuit. | cond-mat.mes-hall |
LuFe 2 O 4 exhibits multiple phase transitions. 2D charge correlations are observed below 500 K, while below 320 K 3D charge order is established, roughly coinciding with the onset of ferroelectricity [2,8]. Magnetic order appears below 240 K and 2D ferrimagnetic order has been suggested by neutron scattering studies [9]. However, strong sample dependent behavior observed in other members of RFe 2 O 4 [7] suggests that unraveling the interesting behavior of LuFe 2 O 4 requires paying due attention to sample quality. In this letter we present extensive neutron diffraction measurements from 20 to 300 K on high quality single crystals of LuFe 2 O 4 . We report several new findings that provide information about the underlying magnetic interactions. | cond-mat.str-el |
The T > 0 dynamical susceptibility has been studied previously using exact diagonalization of finite length chains. 10 We believe our approach provides a useful complement to this work. The numerical approach yields results for all α and is not restricted to small temperatures. However, the system size that can be studied is quite small. We, on the other hand, must proceed perturbatively in α and are restricted to low temperatures, but our calculations do not suffer from finite-size effects. Moreover the nature of low-lying excitations is more apparent and we can identify the specific processes that give rise to the various finite-temperature effects in the structure factor. | cond-mat.str-el |
As R is gradually changed from 0 to 1, Ψ 0 changes from a single-peaked WP at t = 0 to a doubled peaked one at t = T , namely a splitting of a WP occurs. The final time of the fast-forward T F is related to T as T F = εT /v. θ is obtained by numerical integration of equation (2.18) under the boundary conditions ∂θ ∂x (x = 0) = 0 and θ(x = 0) = 0. The spatio-temporal dependence of θ is shown in figure 6. The driving potential is calculated from equation (2.28) with use of equations (3.15) and (3.17) and θ. V F F (solid line) and |Ψ F F | 2 (broken line) which is accelerated by V F F are shown in figure 7. The conversion from a hump to a hollow in the central region of the potential in figure 7 is caused by the deceleration of α(t)ε which suppresses the splitting force. | cond-mat.mes-hall |
Table I shows there is a general correspondence between the decay rate of known correlations and that of the singular values; the degree of correlation in Fig. 2 tends to be overestimated due to the very small range of r. The rung-fermion case (b) at filling 1/4 breaks translational symmetry, with period-2 long-range order. Examination of Fig. 2 (b) indeed shows the corresponding contrast with the other two cases: the singular value for the order-parameter operator (CDW+) is non-decaying and saturates the bound σ = 1/2, whereas other kinds of singular values are orders of magnitude smaller. In the boson-pair case (c), as t ′ grows large (the boson-pair limit), a crossover is expected to asymptotic superconducting (SC) correlations; but Fig. 2(c) shows that CDW correlations still dominate at all accessible distances, similar to hardcore bosons [17]. | cond-mat.str-el |
1. The measurements were performed at a base temperature of T = 40 mK. The pumping frequency was chosen to be f = 50 MHz at a rf power of P RF = -16 dBm. As shown in Fig. 1(a) the pumped current increases in steps of ef as V 2 is made more positive. The red and blue curve compare the dependence of the pumped dc current on V 2 without and with perpendicular magnetic field of B = 3 T applied, respectively. | cond-mat.mes-hall |
45 In the case v 2 R = v 2 D , as already mentioned, there exists a fixed precession axis directed at the angle of π/4 (or -π/4) in the quantum well plane. Therefore, not only the spin current, but also the induced spin density goes to zero in this case. This behavior is demonstrated in Figs. 123, where the calculated real parts of the components of spin polarizability and spin conductivity tensors, Reχ αβ and ReΣ z αβ , are plotted as functions of the ratio of Rashba and Dresselhaus velocities. The case of degenerate electron gas is assumed. The spin polarizability is expressed in the units of static polarizability for symmetric [001]-grown quantum wells, χ 0 = emv D /2πh 2 ν. | cond-mat.mes-hall |
The values of |δρ g xx (B)/ρ 0 A(T /T g )| at the extrema are plotted in Fig. 6 (b) for the three major components, f 2 , f 3 , and f 4 . As can be seen, the function C/ sinh(B w /B) describes the behavior of the normalized amplitudes quite well. Fitting by the two parameters C, B w can be carried out without difficulty for the components with large enough amplitudes, such as f 3 and f 4 in the present case. The scattering parameter µ w =π/B w is expected not to vary among components. This is found to be the case for f 3 and f 4 ; the values obtained by the fitting, µ w =10.2 and 11.3 m 2 /Vs, respectively, roughly coincide (within experimental error ± ∼0.5 m 2 /Vs). | cond-mat.mes-hall |
Whether the transition becomes first order as T → 0 remains to be studied. On increasing µ from 0 the transition occurs at ∆ µ ∆ qp . (At βt = 10, ∆ µ ≃ 3.6t and ∆ qp = 3.8t.) IV. OPTICAL CONDUCTIVITY The optical conductivity can be computed using the Kubo formula and the minimal coupling ansatz p → p -A. The dissipative part of the conductivity is then 21 σ(Ω) = 2e 2 ∞ -∞ dω π d 2 p (2π) 2 f (ω) -f (ω + Ω) Ω ×Tr [j(p)ImG(ω + Ω, p)j(p)ImG(ω, p)] ,(23) where the current operator is j = δH/δp x . | cond-mat.str-el |
The plots are for symmetric source and drain tunnel barriers, and varying core edge asymmetry. We assume a 6 meV bias symmetric about the Fermi energy, and two transport channels at energies epsilon Fermi ± 1 meV. Fock-space coherence between states differing by either one (middle submatrix) or two (bottom submatrix) particles. The remarkable structure evident in Fig. 3 implies that each block undergoes its own independent time evolution. In turn, this suggests that, should a Fock-space coherence be established in the system at any time, this coherence can be robust and long-lived even in the rather severe perturbation of transport carriers being continually injected and removed from the system. | cond-mat.mes-hall |
The energy difference between the ground state and the first excited states is typically 10 -4 -10 -3 eV. This might mean the existence of very low-lying excited states in the limit of large N n d cut . C. Short range correlations Next, we calculate the spin-spin correlation function S(q) for the sixteen sites in the unit of four tetrahedra. Here, S(q) is an equal-time correlation, i.e., a frequency integrated quantity, and defined by 16) S(q) = 1 N s ijg g|S z i S z j |g N g exp(iq • (x i -x j )), (S z i = 1 2 σα σn iασ ,(17) where |g and N g means the index and degeneracy of ground states, respectively. N s is the number of lattice sites (N s = 16 in the present case) and x i is the position of site i. Note that S z i is the spin operator not of a tetrahedron unit but at the vanadium site i. | cond-mat.str-el |
(A2) and (A3) self-consistently. On the other hand, the chemical potential µ 0 of the uniformly doped wire can be found from the particle density of n lm = -g sv ℑG R lm:lm (E)/πa together with the Poisson's equation. The self-energy of Eq. ( 12) caused by the coupling of the device to the source and drain regions is obtained by solving the uniformly doped wire with vanishing boundary conditions. In the similar way to Eq. (A2), it is given by,. | cond-mat.mes-hall |
The inset shows the value of the minimum conductivity and compares with the experimental results. These are the same three samples that were shown in Ref. [4] to compare the conductivity at high carrier density (far from the Dirac point) with a numerical Boltzmann theory, and here we show the low carrier density comparison near the Dirac point. Through Eq. 6, we have the high-density measurements directly giving n imp , and this is the only parameter used to determine the minimum conductivity σ 0 . Our results show that contrary to common perception, the graphene minimum conductivity is not universal, but that future cleaner samples will have higher values of σ 0 . | cond-mat.mes-hall |
The temperature dependence in the experimentally obtained spectra was simulated with the temperature (given in units of eV) used for the Boltzmann averaging over excited states as the only extra parameter, as shown in Fig. 3. All other simulation parameters were fixed to the values given in Table I. In this way we found the expectation values for |S z | at each corresponding experimental temperature, arising from the thermal population of the m S levels in the S = 5/2 multiplet. The temperature dependence of |S z | (Fig. 6) shows a gradual alignment of the spins into the kagome planes, at temperatures well above the transition temperature to a long-range ordered state at 64 K. This is in rough agreement with the degree of co-planarity of the spins as measured using neutron diffraction on large single crystals 13 . | cond-mat.str-el |
This presentation allows for an easier discussion of the results. In figure 2 we have plotted the total spin moment in the unit cell in µ B within the CPA approach. If the alloys under study are half-metals they should follow the Salter-Pauling behavior for the total spin moments, M t : M t =Z t -24 [20]. The total spin moment in µ B is just the number of uncompensated spins and thus the number "24" arises from the 12 occupied minority spin states (for details see reference [20]). Z t denotes the total mean number of valence electron and is given by the expression plotted the atom resolved spin moments for the transition metal atoms Co, Cr and Mn, respectively, within the CPA approximation. Z t = 2 * z Co + (1-x) * z Cr +x * z Mn +(1- In each figure we present the moment as a function of the concentration y in the upper panel with different bars corresponding to different x and as a function of x in the lower panel with different bars corresponding to different values of y. | cond-mat.mtrl-sci |
Recently, Meek et al. discussed the electronic properties of uranium dioxide and revealed the potential performance advantages of uranium dioxide as compared to conventional semiconductor materials [7]. Especially, the higher dielectric constant of UO 2 makes it more suitable for making integrated circuits [7]. This may stimulate many studies of the optical properties for actinide dioxides in future. Optical adsorption and reflectance spectra of semiconductors have been studied for several decades both experimentally and theoretically, whereas, similar works performed on actinide dioxides is still very scarce although they are necessary not only from the viewpoint of basic science but also from their technological importance in industries. Experimentally, Schoenes studied the incidence reflectivity of UO 2 single crystals in the photon energy range of 0.03-13 eV, from which the complex dielectric function ε(ω) = ε 1 (ω) + iε 2 (ω) has been derived [8]. | cond-mat.str-el |
It is brought about by the distortion of the orientation of the magnetic moment within the basal plane. In neutron scattering, the signals at the (6 ± 1)th magnetic satellite spots were observed. 1 In the present case of RXS in the E1 transition, the intensities at the (6 ± 2)th as well as (6 ± 1)th order satellite spots, which are related to rank two and one operators, respectively, will be detected. By evaluating the Ho case in the vicinity of the M 5 absorption edge, we can conclude that the intensity at the fourth satellite spot is within the reach of the present experimental condition of the E1 transition in the π -σ ′ channel. As for the numerical results, first, we have concentrated on the comparison of our result with the experiment in the uniform helical phase. The calculated RXS spectrum at the first satellite spot and the absorption coefficient show excellent agreement with the observed ones both the spectral shape and the peak position (for the former spectrum). | cond-mat.str-el |
The third rank tensor χ lmn describing the linear PGE is symmetrical with respect to the interchange of the second and third indices, and γ lm is a second-rank pseudotensor describing the circular PGE. The (001)-grown heterostructure has the point-group symmetry F ref = C 2v which forbids in-plane photocurrents under normal incidence 19 as it is confirmed by measurements performed on the reference samples. The lateral superlattice can reduce the symmetry of the system. Let us denote the superimposed periodic lateral potential as V (ρ), where ρ is the two-dimensional radius-vector, and introduce the period a and two inplane axes x and y oriented, respectively, parallel and perpendicular to the direction of periodicity. Then, by definition, the lateral superlattice potential does not depend on y and is a periodic function of x, namely, V (x + a) = V (x). One of the symmetry elements of this potential is the mirror reflection plane σ y perpendicular to the axis y. | cond-mat.mes-hall |
The wall is assumed to be rigid and one-dimensional, described by two collective coordinates, position X and angle out of easy-plane φ 0 . Spin-transfer torque arising from angular momentum conservation was shown to contribute to wall velocity, and spin relaxation and non-adiabaticity were shown to work as a force on the wall, which induces φ0 . Solving the equation of motion, we found that there is a threshold current to drive the wall arising from hard axis magnetic anisotropy energy K ⊥ and/or extrinsic pinning potential V 0 . Threshold current is determined by K ⊥ in the intrinsic pinning regime, by V 0 , K ⊥ and force from the current if in extrinsic pinning regime. Our Acknowledgement The authors are grateful to Y. Yamaguchi, T. Ono, M. Yamanouchi, H. Ohno, Y. Otani, H. Miyajima, M. Kläui, Y. Nakatani, A. Thiaville, E. Saitoh, K.-J. Lee, A. Brataas, R. Egger, M. Thorwart, J. Ieda, J. Inoue, S. Maekawa and H. Fukuyama for valuable discussion. | cond-mat.mes-hall |
In general, two mechanisms are known to generate helicoidal-type structures; competition between nearest and next-nearest neighbor interaction and direct DM interaction, resulting from the loss of a center of symmetry relating magnetic ions. It is easy to show that in none of the magnetic structures described here does the cycloidal modulation lower the next-near neighbor magnetic energy, regardless of the sign of the interaction. This suggests that the observed c-axis modulation may be the effect, rather than the cause, of the loss of the center of symmetry relating Mn 4+ ions. V. SUMMARY The magnetic structures in the ferroelectric/commensurate magnetic regime of three com- It also contains zig-zag AFM chains, in disagreement with what was previously reported [21]. This comparative study strongly supports symmetric exchange as the principal mechanism leading to ferroelectricity since the small non-collinearity of the magnetic moments within chains along c is observed only for the YMn 2 O 5 and HoMn 2 O 5 compounds. This does not preclude a potential, but much weaker contribution from Dzyaloshinskii-Moriya interactions. | cond-mat.mtrl-sci |
For each P site the line position was found to shift with temperature. However for the strongly coupled P1 site the shift was found to be much stronger compared to the weakly coupled P2 site and is also strongly orientation dependent. In the present work we have mainly focused on the P1 site and therefore data for the P2 site are not shown. Figure 3 shows the 31 P NMR spectra for the P1 site measured on a single crystal at different temperatures and field applied along different orientations. Figure 4(a) presents the temperature dependence of K for the P1 site derived from the data in Fig. hand for the P2 site the shift is very weak. | cond-mat.str-el |
As discussed earlier, the experiments suggest a term in G that breaks the symmetry P → -P ; this is achieved by linear coupling of P to an external electric field E ext and/or to an effective bias field 10 which we take to be of the formW l = W 0 e -l/lw ,(2) where l w ∼ l c . We note that the thickness-dependence of W l is included to model the increased smearing of the dielectric susceptibility with decreasing l of ferroelectric films. 10 At present we will treat W l phenomenologically, and will defer discussion of its exponential decay and its possible origins to Section V. Putting all these elements together, we begin our phenomenological study with the free-energy expansion G(P, σ, T ) = 1 2 α(T )P 2 + 1 4 γP 4 -(W l + E ext )P -Q 11 σ zz P 2 -Q 12 (σ xx + σ yy )P 2 - 1 2 s 11 (σ 2 xx + σ 2 yy + σ 2 zz ) -s 12 (σ xx σ yy ) -s 12 σ zz (σ xx + σ yy ) - 1 2 s 44 σ 2 xy ,(3) where α(T ) = α(T ) + α d ; α(T ) = β(T -T bulk ), T bulk is the bulk transition temperature, α d is discussed below, and β and γ are Landau coefficients; here Q ij and s ij are the electrostrictive constants and the elastic compliances at constant polarization respectively. The depolarization field contributes to the free energy through the coefficient α(T ) in Eq. (3) 26,34α d = l e ǫ 0 ǫ e l ,(4) where l e is the screening length of the electrodes, and ǫ 0 and ǫ e are the electric permittivities of the vacuum and the electrodes respectively. The mechanical conditions in the film are ∂G/∂σ xx = ∂G/∂σ yy = -u l , ∂G/∂σ xy = 0 and ∂G/∂σ zz = -u zz 4 . | cond-mat.mtrl-sci |
It too is seen to increase linearly with field, implying a lowering of effective mass with an increase in field; and which behavior is consistent with a similar finding for the single impurity Anderson model 18 . The field-dependence of the full density of states is illustrated in figure 6, where we plot the (universal, strong coupling) conduction band density of states D c (ω; h), as a function of the ω, for various h eff . The solid curve h eff = 0 represents the insulating ground state, while the dotted curve is for h eff = 0.38, which is just above the insulator-metal transition, so the gap has closed. The remaining curves are for h eff = 1 and h eff = 5, showing metallic densities of states characterised by a finite spectral density D c (0; h) at the Fermi level. In the non-interacting limit, U = 0, the spectral gap closes linearly with the applied field as in equation 12, and the essential mechanism for the insulator-metal transition is obvious: Zeeman splitting moves the up-and down-spin bands rigidly, resulting in their crossing at a critical field, h c0 /∆ 0 (0) = 1 2 . This simple picture is naturally modified in the presence of correlations, U > 0, where two essentially competing effects are operative. | cond-mat.str-el |
The LS of the closed system can be viewed as a superposition of two counterpropagating waves in the open system coming from the left and right lead. These waves obey a certain phase relation to fulfill the boundary conditions at the outermost walls of the lead stubs which can be extracted from the eigenstates. For δ = 1.5, only the (n, 1) states are of the LS type. They cause a large transmission that is only modified by the sharp resonances due to the CS. For δ = 1.25 (Fig. 1.b) the resonances are caused by (n, 3) states with n = 4, . | cond-mat.mes-hall |
The results obtained by the two different pressure setups agreed reasonably in the temperature range of overlap, indicating a very good homogeneity of pressure. Pressure was calibrated by measuring the superconducting transition temperature of Pb [13]. The inset of Fig. 1 demonstrates ρ(T ) of NiS 2 at relatively low pressures below 4 GPa. With applying pressure, the insulating behavior of ρ(T ) switches into metallic behavior, indicating the occurrence of metal-insulator transition. In between 2.6-3.4 GPa, we observe a discontinuous jump of resistivity as a function of temperature, which corresponds to a first order metal-insulator transition line on the phase diagram in Fig. | cond-mat.str-el |
The momentum and displacement polarization vectors are related by, e.g., µ = -iωǫ + Aǫ. We can verify that the canonical commutation relations are satisfied, [u l , p T l ′ ] = ihδ l,l ′ I, andH = k hω k (a † k a k + 1/2). Based on a definition of the local energy density and the continuity equation for energy conservation, an energy current density can be defined as [6,7,11], J c = 1 2V l,l ′ (R c l -R c l ′ )u T l K l,l ′ ul ′ ,(8) where the index c = x, y, or z labels the cartesian axis, V is the total volume of N unit cells. The components of the current density vector can be expressed in terms of the creation/annihilation operators. Ignoring the a a and a † a † terms which vary rapidly with time, one obtains [7] J c = h 4V k,k ′ ω k ω k ′ + ω k ′ ω k ǫ † k ∂D(k) ∂k c ǫ k ′ a † k a k ′ δ k,k ′ . (9) The thermal conductivity tensor can be obtained from the Green-Kubo formula [12], κ ab = V T βh 0 dλ ∞ 0 dt J a (-iλ)J b (t) eq ,(10) where β = 1/(k B T ), the average is over the equilibrium ensemble with Hamiltonian H. The time dependence of the annihilation operator is trivially given by a k (t) = a k e -iω k t . | cond-mat.mes-hall |
Due to the fact that filling the dot with a single electron occurs with rate 4Γ L and adding a second electron only with rate Γ L the singly occupied state |+ L is, in lowest order in (a + 1), four times more likely than an empty dot. Occupation of the triplet is even rarer than singlet occupation: it starts in order (a + 1) 2 , because it can only be reached via singlet occupation and subsequent decay to the state |-R . It is eminent from the flowchart Fig. 9 that there are several distinct cycles through which electrons are transported from left to right: the transitions |+ L ↔ |S and |0 → |+ L → |S → |-R and then back to |0 , or several sub-cycles via |T . As these cycles transfer electrons at different mean currents and with different statistics, it is clear that a complicated telegraph effect will lead to increased noise. In contrast to the channel exclusion described in the previous sections this effect is not related to separated Hilbert spaces. | cond-mat.mes-hall |
Henceforth we shall restrict the meaning "quasiparticle" to those whose ℓ ℓ ℓ-vector is an integer combination of ℓ ℓ ℓ 1 through ℓ ℓ ℓ 7 . From K -1 r it can be seen that the system contains quasiparticles with non-trivial mutual statistics. In particular, there there are fermions having semionic mutual statistics (i.e., a phase factor of π when one quasiparticle winds around another), manifesting in, e.g., quasiparticles described by ℓ ℓ ℓ 4 and ℓ ℓ ℓ 5 . The self-statistics and mutual statistics of different quasiparticles can be understood intuitively. Recall that our system is constructed by coupling integer and fractional quantum Hall states via a common constraint gauge field α. If we assume that the different quantum Hall states are independent of each other, i.e., a "charge" in one matter-field component has trivial bosonic statistics with a "charge" in a different matter-field component, then the statistics of these quasiparticles can be read off by considering their underlying constituents. | cond-mat.str-el |
Notable inter-octahedral structural parameters are the (Fe-Fe) short and (Fe-Fe) long distances, which decrease and increase, respectively, at the spin transition as can be seen from Table I. The same was observed by Rollman et.al. 16 in their GGA calculations. Although also the remote surroundings (i.e. the next to next-nearest neighbors and inter-octahedral distances) have non-negligible effects on the nature of the spin transition, the most significant changes observed at the transition concern mainly the octahedra, as already pointed out by a number of studies. 25,26,27,28,29,30,31 Of the intraoctahedral structural parameters, the average Fe-O bond lengths, as shown in Fig. | cond-mat.mtrl-sci |
Near the crossing, two poles of the Green's functions dominate the transmittance as discussed in Section II C, and they contribute with opposite sign to it. As a consequence a strong depression of the conductance takes place at φ d . In particular, Eqs. ( 5) and (6) indicate that the transmittance vanishes at φ = φ d for the value of the gate voltage at which the energies of both 5-electron states coincide with that of the ground state for 6 electrons. This might be an artifact of these expressions which are perturbative and are not expected to be valid near this point of triple degeneracy. 36 However, the physical origin of the depression of the conductance is clear and should be present in a more elaborate treatment. | cond-mat.str-el |
The dependence on the Coulomb interactions, i.e., on the ratio d/a e is also negligible: as it is reduced from 0.8 to 0.2 the oscillator strength reduces continuously from 1.4912M 2 to 1.4780M 2 . As these parameters change the sum of the two oscillator strengths of transitions A and B remains a constant equal to 3/2M 2 . This is evidence that our calculations are correct. The highest peak with the energy 6.61E C has the optical strength 5/2M 2 and originates from the transitions |0, 1 2 , -1 2 1 → |0, 1, -1 1 and |Φ 1 , or |0, -1 2 , 1 2 1 → |0, -1, 1 1 and |Φ 1 , see Fig. 5. In these transitions there are two possible final states with the same energy, and the oscillator strength is the sum of the two contributions. | cond-mat.mes-hall |
In Sec. II, the model of Ref. [44] is developed to include the AB-flux. In Sec. III, we discuss the CBSIP in the presence of the AB-flux both for a single and for a couple of impurities. II. | cond-mat.mes-hall |
The Wigner-Smith matrix is N = (2πi) -1 s † ds/dE. We introduce a local density of states matrix in which the energy derivative is replaced by a functional derivative with respect to the electrostatic potential U (r), n(r) = -(2πi) -1 s † δs/δeU (r). Its trace is the local density of states ν(r) = T r[n(r)] = -(2πi) -1 T r[s † δs/δeU (r)]. Coulomb interaction leads to a screened charge described by an effective density of states and an effecitve density of states matrix [33] Here g(r, r ′ ) is a an effective interaction. It is the Green's function of the Laplace equation with a non-local screening kernel π(r, r ′ ), and determines the solution of the potential (matrix) u(r) = d 3 r ′ g(r, r ′ )n(r ′ ). In terms of the effective density of states matrix the capacitance of the structure is The charge relaxation resistance is T r[( d 3 r n ef f (r)) † ( d 3 r n ef f (r))] ( | cond-mat.mes-hall |
This is consistent well with the experimental assignment [19] by Naegele et al., who attributed the peak around 3 eV in ε 2 (ω) to intra 5f 2 transitions, while the peak structures above 5 and 10 eV were ascribed to the f → d and p → d transitions, respectively. Another assignment was suggested by Schoenes according to their dielectric function deduced from the reflectivity measurement; they argued that the peaks near 3 and 6 eV correspond to f → d transitions, and that the peaks near 8 and 11 eV are due to p → d transitions [8,20]. Herein, the assignment of f → d transition at 3 eV in Ref. [8,20] is not supported by our calculation. The cause is that in assigning the peak in ε 2 (ω) at 3 eV, the energy distance between U occupied 5f 2 and O 2p valence bands was overestimated in Ref. [8,20] to be as large as 4 eV, which is much larger than that directly determined by the photoemission measurements [16,21,22]. | cond-mat.str-el |
For the former we have calculated the power numerically and derived analytic expressions in low-and high-temperature limits. The power transfer to acoustic phonons dominates the diffusion power above a certain crossover temperature, estimated in Eqs. (32) and (33). At even higher temperatures, there is another crossover where optical phonons begin to dominate, and we have estimated also these crossover temperatures numerically (Fig. 4). We find that for graphene on the substrate the most relevant optical phonons are likely to be the surface optical modes of the substrate. | cond-mat.mes-hall |
Furthermore, from Figure 3A it can be seen that graphene is still accurately quantized at 4.2 K. At this temperature the measurement current has to be reduced to 2.3 μA and consequently gives a larger uncertainty over a comparable time interval. Nevertheless, these measurements demonstrate that in epitaxial graphene is accurately quantized and can be used for precision resistance metrology even at elevated temperatures. R K /2 against the measurement time τ. The data plotted in such way can be accurately fitted by 1/τ 1/2 dependence -the behaviour typical of the predominantly white, i.e., random and uncorrelated, noise. This justifies the use of the standard measures of uncertainty and suggests that even these very accurate results can be further improved if one is prepared to measure for longer. As seen in the AFM images the graphene Hall bars are patterned across many substrate terraces. | cond-mat.mes-hall |
If c is the dilution, we assume that at each site we have a hole with probability c and a spin with probability 1 -c. The holes are fixed in their position and extensive quantities are aver-aged over all possible configurations C using the relationO = C P (C)O(C),(7)whereP (C) = c N h (1 -c) Ns(8) is the probability of the configuration C with N h holes and N s spins. The entropy and specific heat for several values of hole concentration are shown in Fig. 12. We note that at these intermediate temperatures [Fig. 12(a)], holes simply lower the entropy at all temperatures. In the case of the specific heat, they just displace the high temperature peak to lower temperatures, almost without changing its E. Entropy difference from kagome lattice Heisenberg model and implication for ZnCu3(OH)6Cl2 To conclude the section on specific heat and entropy, we study the entropy difference between the KLHM and the model with different DM parameters. | cond-mat.str-el |
Minor adaptations are necessary for the XX model [17]. The orbitals available for occupation by spinons are equally spaced at ∆κ = 2π/N and their number is (N -N s )/2+1, where N s = 0, 2, . . . , N for even N and N s = 1, 3, . . | cond-mat.str-el |
In the case of the quantum Hall (QH) effect, it is the quantization of the Hall conductance and the fractional charge of the elementary excitations which are a result of non-trivial topological structure. The Z 2 topological invariant gives a correct mathematical characterization of the QSH state [10]; however, unlike the TKNN quantum numbers [11] of the quantum Hall (QH) state, it is not directly measurable experimentally. In this work, we show that for the QSH state a magnetic domain wall induces an elementary excitation with half the charge of an electron. We also show that a rotating magnetic field can induce a quantized dc electric current, and vice versa. Both of these physical phenomena are direct and experimentally observable consequences of the non-trivial topology of the QSH state. The idea of fractional charge induced at a domain wall goes back to the Su-Schrieffer-Heeger (SSH) model [12]. | cond-mat.mes-hall |
We have done the NMR measurements at two different radio frequencies of 70 MHz and 11.13 MHz which correspond to an applied field of about 4.06 T and 0.65 T, respectively. Spectra were obtained either by Fourier transform of the NMR echo signals or by sweeping the field. The NMR shift K = (ν -ν ref )/ν ref was determined by measuring the resonance frequency of the sample (ν) with respect to nonmagnetic reference H 3 PO 4 (resonance frequency ν ref ). The 31 P spin-lattice relaxation rate 1/T 1 was measured by the conventional single saturation pulse method. For the analysis of NMR data, we measured the magnetic susceptibility χ(T ) of the single crystal at 4 T for field applied along c-axis in a commercial (Quantum Design) SQUID (Superconducting Quantum Interference Device) magnetometer. III. | cond-mat.str-el |
The effective anisotropy per layer is essentially the same as for a single layer, thus the critical S z -value (magnetization at critical field B crit ) is practically the same. The critical temperature is higher than that of a monolayer due to the increased magnetic stiffness of the double layer. IV. SUMMARY AND CONCLUSIONS Using GFT and QMC calculations we studied easyplane systems as well as easy-axis systems with an external field applied perpendicularly to the film. The GFT treatment of the Hamiltonian Eq. ( 1) consists of a RPAdecoupling for the nonlocal terms and an AC-decoupling for the local terms performed in a rotated frame, where the new z ′ -axis is parallel to the magnetization. | cond-mat.str-el |
However, as seen in this work, again these modifications do not alter the qualitative morphological implications of the continuum theories to be described below, which justifies (at least for temperatures at which Ehrlich-Schwoebel barriers do not dominate surface diffusion [177]) the strong similarities in nanopattern formation by IBS on metals, as compared with amorphizable targets. As a morphological theory, again Sigmund's [154] has some limitations: it does not predict the alteration of the morphology during the process for scales much larger than penetration depth for ions; as surface diffusion is not considered, the wavelength of the pattern needs to be of the order of the length scales appearing in the energy distribution Eq. ( 5); the effects of surface shadowing or redeposition are not considered; it does not predict the time evolution of the morphology and how it affects the rate of erosion. Thus, an additional physical mechanisms and a more detailed description of the surface height are needed in order to derive a morphological theory with an increased predictive power. Monte Carlo type models (Kinetic) MC methods have a long and successful tradition in the context of IBS [200,201]. For our morphological purposes, they have found increased application during the last decade. | cond-mat.mtrl-sci |
II. THEORETICAL FORMULATION The Keldysh formalism had been applied previously by Edwards et al. 11 to calculate the spin-transfer torque in the CPP geometry. An essential requirement for the implementation of the Keldysh formalism is that a sample with an applied bias can be cleaved into two noninteracting left (L) and right (R) parts by passing a cleavage plane between two neighboring atomic planes. It follows that, initially, neither charge nor spin current flows in the cleaved system although the left and right parts of the sample have different chemical potentials. This is most easily achieved for a tight-binding (T.-B.) | cond-mat.mtrl-sci |
Solution-processable organic photovoltaic cells have shown a promising performance increase in the recent years 1 . A major focus now lies in finding a viable strategy for a further optimization of the power conversion efficiency, guided by a deeper understanding of the fundamental processes. Device simulations are useful tools to assist in finding such routes, as they allow the extrapolation of possible but not yet implemented device concepts. We present macroscopic simulations of polymer-fullerene solar cells based on an effective medium approach. The influence of the charge carrier mobility on dissociation and transport processes, which are governing the power conversion efficiency, will be covered. Based on these results, we will discuss the most promising optimization routes for organic solar cells. | cond-mat.mtrl-sci |
I. INTRODUCTION In recent years, a lot of work has been conducted on the interaction of alkanethiols with gold surface because of their ability to form self-assembled monolayers (SAMs); see Ref. [1] for a review. These SAM films have wide application on wetting phenomena, tribology, chemical and biological sensing, optics and nanotechnology. The consistent model for the methanethiol adsorption process on the Au(111) substrate was proposed recently [2,3]. It was revealed that the methanethiols stay intact on the regular Au(111) surface, but the S-H bond ruptures on the defected Au(111) surface. | cond-mat.mtrl-sci |
Care should be taken with off-diagonal operators. will be published elsewhere [19]. In our simulations, this procedure reduced the computational effort for measuring Greens functions in all of momentum space by about a factor of 5N , where N is the system size. We improved the inversion of ( 2) with the maximum entropy technique (MaxEnt) [25] by using the exactly known moments M 1 = ω0 2 and M 3 = ω0 3 2 , whereM n = ∞ -∞ dωS (q, ω) ω n . These moments roughly double the resolution of S (q, ω) near ω ∼ ω 0 . The temperature in our simulations was chosen low enough to achieve the T = 0 limit where no changes in S (q, ω) by lowering the temperature are observed. | cond-mat.str-el |
( 21) and calculated with the use of the reduced inverse susceptibility yQ(T /T0) shown in Fig. 1. First, we analyze the lowest-order approximation to the low temperature (T ≪ T * ) resistivity, ρ(T ) ≈ ρ imp + ρ (1) sf (T ),(22) where ρ (1) sf (T ) = A (1) sf f (T /T 0 ),(23) and compare its T -dependence with that of the observed 10 experimental resistivity. Here A sf = (2πcT 0 /T A )X -2 1 C 1x,1x(1) is an adjustable parameter (together with ρ imp ) in a low-T fit procedure using the calculated f (T /T 0 ) shown in Fig. 2. One may see that the function f (T /T 0 ) nearly precisely follows the quadratic dependence, f (T /T 0 ) = c 1 T 2 with c 1 = 0.0033, for T < 2K ≪ T * , where the Fermi-liquid behavior [ρ exp (T ) -ρ imp ] = AT 2 in LiV 2 O 4 was reported. | cond-mat.str-el |
. . Note that H eff CORE reproduces exactly the low-energy physics if one considers all of the terms on the right-hand side. In practice, it is necessary to perform a truncation. The convergence of the algorithm therefore depends both on the range of the operators taken in the cluster expansion and on the number and type of low-energy states retained for one subunit. Hence the successful application of the CORE technique does require some physical insight concerning the problem at hand. | cond-mat.str-el |
( 25) if the second-and thirdorder terms differ from those defined by a Hydrogen-like density expansion, e.g., Eq. ( 26). Thus, the foregoing cancellation fails for a density with power series expansion n(r) ∝ 1 -2Zr -(4/3)Z 3 r 3 + . . .. This fact will limit applications of simple κ 4 -based KE functionals to those densities which have precisely Hydrogen-like behavior up to fourth order. | cond-mat.mtrl-sci |
The effect of pressure on the resistivity and functional properties of CMR manganites has been studied widely in the past. In particular, resistivity versus pressure has been measured in La 0.88 Sr 0.12 MnO 3 thin films where it was found that resistivity decreases up to 0.81GPa, after which it increases rapidly. Here, a continuous and almost linear decrease of the resistivity is observed up to 2.5 GPa. The combined effect of magnetic field (0,2,5,7 and 9T) and high pressure (0-2.5GPa) on the resistivity of Pr 0.5 Ca 0.5 MnO 3 is presented on Figure 2. Data are presented for the warming run and no thermal hysteresis was observed. Without magnetic field, the film is semiconducting, while a 7T applied magnetic field (not shown) induces a metal-to-insulator transition (MIT) around 150K, which leads to a large magnetoresistance value. | cond-mat.mtrl-sci |
The first example is for a machine mount to reduce acoustic vibration from engines and machines. The second example is frequency selective filtering in waveguides, and an acoustic lens. From the above we see that phononic crystals are obviously useful for blocking sound at specific frequencies. Using them for acoustic signature reduction in machinery is an obvious application. Conventional machine mounts are essentially heavy springs. They typically have a flat performance between 500 Hz and 10 kHz and dampen signatures by about 20-30 dB. | cond-mat.mtrl-sci |
High-pressure reflection high-energy electron diffraction (RHEED) was used to monitor the growth of the SRO film in-situ. By monitoring RHEED oscillations, SRO growth was determined to proceed initially in a layer-by-layer mode before transitioning to a step-flow mode. RHEED patterns and atomic force microscopy imaging confirmed the presence of pristine surfaces consisting of atomically flat terraces separated by a single unit cell step ( 3.93 Å). X-ray diffraction indicated fully epitaxial films and x-ray reflectometry was used to verify film thickness. Bulk magnetization measurements using a SQUID magnetometer indicated a Curie temperature, T C , of ∼ 150K. Sensitive detection of FMR by the time-resolved magnetooptic Kerr effect (TRMOKE) has been demonstrated previously [12,13,14]. | cond-mat.mtrl-sci |
In electronic transport with current-flow perpen-dicular to the layer planes (CPP geometry) of a metallic multilayer, the interface specific resistance AR (area A through which the CPP-current flows times the sample resistance R) is a fundamental quantity. In the past few years, measurements of AR have been published for a range of metal pairs [1][2][3][4][5][6][7][8][9][10][11][12][13][14][15]. Special interest focuses upon pairs M1 and M2 that have the same crystal structure and closely the same lattice parameters a o -i.e., Δa/a o ≤ 1%, since AR for such pairs can be calculated with no free parameters. That is, taking a given crystal structure and a common a o as known, the electronic structures for M1 and M2 can be calculated without adjustment using the local density approximation, and then AR M1/M2 can be calculated without adjustment using a modified Landauer formula for either interfaces that are perfectly flat and not intermixed (perfect interfaces), or for interfaces composed of two or more monolayers (ML) of a 50%-50% random alloy (50-50 alloy) [16][17][18][19][20][21]. For all four such pairs (Ag/Au [18,20], Co/Cu [18,20,21], Fe/Cr [17,18], and Pd/Pt [12]) where experimental values of 2AR M1/M2 have been published, Table I shows that the previously calculated values for perfect and 2 ML thick alloyed interfaces of these pairs are not very different, and that the experimental values are generally consistent with both values to within mutual uncertainties. In contrast, when Δa/a o is ~ 10%, the agreement between experiment and theory is only semi-quantitativeexperiment and calculations differ by amounts as low as 50% to more than factors of two [11]. | cond-mat.mtrl-sci |
( 6) with k0 = 10 5 cm -1 , d = 10 -6 cm, a = 0.1 at t=15 (in units of t0). for hole probability density the space-time evolution of the spin density strongly depends on the parameter a. In particular, for small values of a as was demonstrated in Sec.II the splitting of wave packet is absent and oscillations of the probability density (ripples) arises on its periphery. Similar behavior of spin density S z ( r, t) for a = 0.1 one can observe in Fig. 5. Using Eq. | cond-mat.mes-hall |