sentences
stringlengths 11
4.62k
| source
stringclasses 3
values |
|---|---|
Calculations were carried in a cubic supercell of 10 Å lateral side for the smaller cube and 20 Å for the larger one. We sampled the supercell BZ at the Γ point only. For the neutral nanocubes, we verified that the residual interaction between periodic replicas is negligible. However, for the charged cubes, the residual interaction is not negligible and was removed by the technique described in Ref. [22]. Results We will begin this Section by describing the electronic states of the infinite ideal neutral NaCl(100) surface, followed by results for charged NaCl(100), where the extra electron and the extra hole are fully delocalized. | cond-mat.mtrl-sci |
(36) In this case, the conductance decays (at fixed R 1 ) with the characteristic length L ≈ R 2 as G ∝ L -π/ϑ . The reciprocal decay, observed in Section III for the full disk, now appears at ϑ = π. As a next example, we consider the conformal transformation, which changes the quantum dot shown in Fig. 1(c) into a rectangle of the width W and the length L. The transformation is given by the formula [25] z -z 0 = W ϑ ∞ Log w + r w -r ,(37)with the condition (R 2 -R 1 + r) 2 /(R 2 -R 1 -r) 2 = e ϑ∞L/W , which leads to Λ(R 1 , R 2 , ϑ) = r -R 1 + R 2 r + R 1 -R 2 2π/ϑ∞ . (38) The origin of the coordinate system of Fig. 2 is now shifted to z 0 ≡ (L + iW)/2. | cond-mat.mes-hall |
Lacking detailed information on the phonon spectrum of SWNH, an assignment of the vibrational frequency ħω c ~ 1.3 meV must remain tentative. This frequency is certainly below the phonon spectrum of molecular hydrogen [20] so it must be related to low-energy modes of the SWNH substrate. In spite of its relevance to explain our experimental findings quantitatively, we note that the role of carbon-substrate motions has been largely neglected in simulation work to date [7,9,10]. Also, since ħω c ~ V 2 , the coupling between internal (rotational) and center-of-mass (translational) degrees of freedom is likely to play an important role in dictating the precise nature of H 2 motions on the SWNH substrate. In this context, recent 13 C NMR relaxation data [21] have identified two distinct processes with disparate relaxation times. On the basis of previous information on the electronic subsystem [22], these can be assigned to either fast processes occurring in SWNHs in close proximity to the surface of the bundles or to slow motions taking place inside the graphitic cores. | cond-mat.mtrl-sci |
B. Landau-Lifshitz-Bloch + Slonczewski equation In the second approach, we add a Slonczewski torque term to the LLB equation. To derive the LLB equation, a probability distribution for the spin orientation is assumed, which is used to find the ensemble average of Eq. ( 5). In addition, the nearest neighbor exchange field is replaced by its mean-field value. The details of the derivation follow closely those in Ref. 11, so we omit them here. | cond-mat.mes-hall |
It is well known that an increasing number of atoms in the cluster causes the one-electron DOS to approach the band DOS of the solid. Increasing the number of atoms to simulate a bulk situation enlarges the number of available transitions to the final states to a point where the computational cost is enormous, while no appreciable change in the spectroscopic observables is obtained. In fact, the quasi-atomic nature of Auger emission found in our ab-initio calculations forces the hole-hole DOS to be centered on the atom where the initial core-hole has been created. We have investigated the role of cluster size by carrying out calculations on a smaller cluster of atoms, Si 2 OH 6 , to see how it compares with the bigger cluster. This nanocluster represents the minimal cluster size which still describes the chemical environment found in solid SiO 2 . In Fig. | cond-mat.mtrl-sci |
substituting the finite differences with the differentials), yields: i i i ε ε ε - = Δ / [ ] ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎦ ⎤ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎣ ⎡ + + - + + - + + - + - - + - - + - - + = F A Y F A Y l l l l l l l l A 2 2 2 1 2 2 1 1 2 1 1 2 1 2 1 2 1 2 2 2 2 2 2 2 2 2 2 2 1 1 1 1 1 1 1 1 1 1 sin 0 cos sin sin 0 0 0 sin sin cos sin 0 0 sin 0 cos 1 0 1 0 cos 0 sin 1 0 0 1 0 sin 0 cos 1 1 0 0 cos 0 sin 1 0 1 α α α α α θ α θ α α α θ α α α θ α α ε α α ε α α ε α α ε [ ] ( ) ( ) ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎦ ⎤ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎣ ⎡ + - + = 0 0 0 0 sin 1 1 cos 1 1 1 1 1 1 α ε α ε b , [ ] , [ ] ( ) ( ) ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎦ ⎤ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎣ ⎡ + - + = 0 0 sin 1 1 cos 1 0 0 2 2 2 2 2 α ε α ε b ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎥ ⎦ ⎤ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎢ ⎣ ⎡ = 2 1 2 1 d d d d d d d ε ε α α v u x [ ] [ ] [ ] [ ] ( ) 2 2 1 1 1 d d d l b l b A x + = - Eq. (3b) in the limit of small variations, gives: 2 2 2 2 2 1 2 1 1 1 2 2 2 2 2 1 1 1 1 1 d 2 1 d 2 1 d d d l A Y l A Y l A Y l A Y E ε ε ε ε ε ε + + + =as well as eq. (3c) poses: v F u F W d sin d cos d θ θ + = According to eq. ( 3a) and (3d) the delamination force can now be easily obtained. For example, considering the symmetric case ( α α α = = 2 1 , , l l l = = 2 1 2 π θ =, and consequently 0 = u andε ε ε = = 2 1 ) we find the following solutions: ( ) ( ) l l l l d sin cos cos 1 d cos cos sin 1 cos 1 1 d 2 2 ε α α α α α α ε ε α ε ε - ≅ + + + - = ( ) [ ] l l l l l l l d sin d d cos 1 cos sin ) 1 ( d cos d 1 cos 1 d α ε α ε α α ε ε α α ε α + + - ≅ + + - + = ( ) ( ) l l l v d sin 1 d sin d cos 1 d α ε ε α α α ε + + + + = The previous equations have been linearized in ε . Accordingly, the energy balance is self-consistently written considering terms up to the second power of ε . | cond-mat.mtrl-sci |
With the insertion of Eq. ( 17) into Eq. ( 15), we get F (T ) = 2πT 0 T A d 3 q (2π) 3 ∞ 0 dλn(λ)[n(λ) + 1] λ 2 λ 2 + (2πu q ) 2 = 2πT 0 T A d 3 q (2π) 3 I(u q ). (18) Here, the last equality defines I(u q ), where u q (t) = y Q (t) + q || /q B 2 + b q ⊥ /q B 2 t ,(19) and t = T /T 0 is the reduced temperature. At the next step, the function I(u q ) can be expressed as I(u q ) = u q 2 ψ ′ (u q ) - 1 u q - 1 2u 2 q ,(20) where ψ(u) is the digamma function and ψ ′ (u) = dψ(u)/du. The integration over q in Eq. | cond-mat.str-el |
It is equivalent to the transition from two-qubit state |ab to state [exp(iθ)] ab |ab , where a, b = (0, 1). Nine-qubit Shor encoding and error syndrome measurement It is known that any realistic quantum computational scheme should be protected from quantum errors caused by the environmental decoherence and the external control imperfections. For this purpose, one may use the quantum error correction codes converting individual qubit state into the specific entangled state of several qubits that, after decoding, restores initial single qubit state (see, e.g., the book of Nielsen and Chuang [1] for detail). In this Section we present the algorithm realizing the nine-qubit quantum error correction code developed by Shor [23]. Auxiliary structure for Shor encoding algorithm The encoding procedure requires us to arrange the qubits in such a way that the interactions between them can be switched on/off on demand in controllable manner. Making use of the results of Sec. | cond-mat.mes-hall |
We show in Fig. 18 (a) an example of this assignment. We use simplified notations for the orbital label in T 1 representations hereafter such as xy ≡ (xy + c 1 z)(x 2 -y 2 ) and so on. It should be noted that the same orbital is assigned to different T z -states depending on bond directions, as depicted. We use for the orbital part eight operators, T µ (µ = 1, 2, • • • , and 8). For orbital degrees of freedom, we introduce a representation that depends on the bond direction. | cond-mat.str-el |
After deconvolution with the instrumental resolution function this line broadening could be assigned to a size-effect, which was best simulated with platelet-like shaped crystallites (ca. 400 Å in diameter and surface parallel to [200]). In further structure considerations, based on the Rietveld method, observed weak microstructural effects were included by simultaneous refinements of the Caglioti profile parameters. For the localization of the copper atoms in the lattice, the occupancies of the different Ni sites were checked. As a result of the best Rietveld fit (Fig. 1) 99% of the copper atoms could be located on the Ni1 site. | cond-mat.str-el |
This sample reproduces very well the structure of a real system (with an overall T (r) discordance [13,14] of 4.0% between 1 and 10 Å). It is pertinent to mention that, whereas the above simulated sample is in excellent agreement with the bulk material, the surface of the simulated particle was not specifically treated to achieve relaxation. Therefore, the surface of our sample is equivalent to a virtual cut in the bulk without any perturbation on the coordination and energetics of the individual atoms. this is clearly only a first order approximation to the real surface. The preparation of a more "physically correct" surface large enough for the meaningful simulation of adsorption processes is a lengthy task not devoid of complications. This is in fact the matter of our ongoing work. | cond-mat.mtrl-sci |
This is because the second intermediate states are dominated by the Sm 3+ configuration which has the same 4 f electron number as the ground state but with different values of J. Then the excitation energy in E -H 0 is of the order of spin-orbit splitting, which is especially small in the case of 4 f 5 with merely 0.12 eV. If we put H hyb ∼ 0.3eV and U eff ∼ E -H 0 ∼ 2eV, we can roughly estimate as H 2 hyb /(U eff ∆ SO ) ∼ 0.4. Hence the ratio of hybridization over the excitation energy exceeds unity, and the weight of mixed states with J = 7/2 or higher is not negligible. Figure 6 illustrates the perturbation processes. The Γ 67 ground CEF states have wave functions with J = 5/2: |a± = 5/6| ± 5/2 + 1/6| ∓ 3/2 , (27)|b± = | ± 1/2 , (28) where a and b specify the orbital quantum number, and ± the Kramers partners. | cond-mat.str-el |
Actually, phonon drag effects are well known in the electron-phonon system [24]. Recently, this issue was considered in the spin liquid context with z = 3 gauge fluctuations [20], where coupled quantum Boltzmann equations for spinon and photon distribution functions are derived. It was argued that such coupled transport equations can be decoupled in some cases, where such drag effects are subdominant, compared with fermion contributions. The present formulation differs from the previous approach in the fact that we did not decompose the gauge field as the study of Refs. [19,20], where the low energy gauge field giving rise to divergence is neglected and only high energy gauge fluctuations are taken. Although the vertex-distribution function itself is not well defined because its part corresponding to the scattering rate is divergent at finite temperatures, we found that such decomposition is not necessary because the formal divergence should be cancelled in the last gauge invariant physical expression. | cond-mat.str-el |
The sample is made on a GaAs/AlGaAs heterostructure using local anodic oxidation [19][20][21] and electron beam lithography. To allow electronic transport, the QD is connected to two leads, source and drain, as schematically shown in Fig. 1. A side gate is used to tune the QD potential. Details about the sample preparation and device properties can be found in Ref. [22]. | cond-mat.mes-hall |
For definitiveness we assume that the spins in the CO phase align themselves along the x axis, whereas the magnetic field is applied along the z axis. We use mean field theory to calculate the uniform and staggered magnetizations and derive therefrom the magnetic phase diagram of the J 2 -λ model. The minimal J 2 -λ model in an external magnetic field B contains only onsite and NNN interactions. Therefore the diamond lattice decomposes into two fcc sublattices with J 2 playing the role of the NN exchange interaction within each sublattice. The Hamiltonian is H min = ij J 2 S i • S j + i H i 0 + B i S z i ,(14) where ij represents nearest neighbor sites on an fcc sublattice and H i 0 is defined by Eq. (1c). | cond-mat.str-el |
7, to justify that the steady state in an extended system has been reached, the currents need to be constant both in time and space. From Fig. 7, we see that j i = const is not fulfilled, although the charge flow in and out of the system is constant, apart from the relatively small oscillations discussed before (compare Fig. 2). This suggests that for the time scales reached in our simulations, the interacting region still undergoes a reorganization of charges and local energies. Indeed, from the data of Fig. | cond-mat.str-el |
Indeed, we gave a general and rigorous proof of the relation (claimed by other authors) between the power law convergence of the series and the number of zero multipolar moments of the crystal construction cell. Based on these convergence analyses we derived a gen- eral real space method with an exponential convergence rate, comparable with the Ewald's method. The exponential convergence is reached as a function of the number of canceled multipolar moments in the construction cell. The crystal is indeed constructed using overlapping construction cells with renormalized charges. We derived a general analytical expression of the renormalization factors, for any given number of zero multipolar moments. Finally, we would like to point out that our method warrants continuous and smooth variations of the renormalization factors. | cond-mat.str-el |
( 4) numerically in II and IV, and utilizing continuity of the wave functions on four sublattices at boundaries, one may obtain the transmission coefficients for a bilayer graphene lattice by using the same equations as in the case of monolayer structure. We first calculate the transmission probabilities of charge carriers through monolayer graphene lattice. The results are depicted in Figs. 4 and5. Fig. 4 shows examples of k y dependence of transmission probability for an abrupt potential barrier with height V 0 = 200meV. | cond-mat.mes-hall |
[24] for definition of the geometric average masses) are m c = 0.01m 0 , m hh = 0.49m 0 , m lh = 0.02m 0 , and m so = 0.07m 0 , whereas the experimental values [25] are m c = 0.07m 0 , m hh = 0.53-0.59m 0 , m lh = 0.08m 0 , and m so = 0.13m 0 . The k-dependent modifications in the GGA+U formalism result in an increase of the bandgap and also a weakening of the bonding-antibonding interaction, thereby improving the effective masses. Using the LAPW method with U = 9 eV we obtain m c = 0.04m 0 , m hh = 0.62m 0 , m lh = 0.05m 0 , and m so = 0.12m 0 which agree much better with the measured values. IV. Conclusion To conclude, we have examined a popular and computationally inexpensive method of correcting for the bandgap underestimation in first-principle calculations. For small values of U one finds that the changes in the valence band are confined to the region around the Γ point. | cond-mat.mtrl-sci |
The corresponding spectral function for spin-↑ and spin-↓ is shown in Fig. 6. Note, that now the spectral function is shown for both positive and negative energies. As before, let us consider first the case of J = 0. Due to an effective exchange field originating from the presence of ferromagnetic electrodes, the spin degeneracy of the dot level is lifted. At zero temperature, the magnitude of the splitting due to exchange field, ∆ε d , can be estimated from the formula 6,7,10 ∆ε d = 2pΓ π |ε d | |ε d + U | . | cond-mat.mes-hall |
For nonidentical dots they are no longer symmetric and they are shifted with respect to F = 0 field, but they distinctly preserve their character [compare with the identical dots case of Fig. 10(b)]. In particular, in both avoided crossings marked by rectangles in Fig. 10(d) at the center of the avoided crossing a single energy level with a much stronger recombination probability than the others appears. Also the energy position of the brighter energy level within the triple of interacting energy levels is conserved. In the lowest-energy avoided crossing it is the lowest energy level. | cond-mat.mes-hall |
For the purpose of characterizing ballistic or diffusive behavior, it therefore does not matter whether the pure fermionic density n i (t) or the averaged quantity ñi (t) is used. In what follows, we will present results extracted from the former, unless stated otherwise. We mention, though, that the quantitative difference in the variance extracted from the averaged as compared to the bare density becomes more pronounced the larger ∆ and the smaller B 0 is. This becomes evident in the case of ∆ = 1, included in Fig. 5(a). The key observation from Figs. | cond-mat.str-el |
. . H mix (t 1 ). (47)Because H mix = σ (H dσ mix + H d † σ mix ) with H dσ mix = α=L,R p V α p a α † p,σ d σ , H d † σ mix = (H dσ mix ) † and the time evolution conserves the spin, we need for each σ separately an equal number of creation and annihilation operators on the Keldysh contour 0 → t → 0: 1 = mσ +nσ=m ′ σ +n ′ σ σ i mσ+m ′ σ (-i) nσ+n ′ σ × t 0 d tσ 1 . . . | cond-mat.mes-hall |
2, and to analyze the CO amplitude of each Fourier content thus singled out. The first step is to subtract the slowlyvarying background to obtain the oscillatory part. In the case of a periodic LSL, the background can readily be defined as the average of the upper and lower envelope curves; 42 since the CO amplitude decreases monotonically with decreasing magnetic field, the envelope curves obtained as spline curves connecting maxima or minima, and hence their average, are also monotonic. A similar approach is not applicable for FLSLs having nonmonotonic envelope curves. We, instead, make use of the numerical differentiation by B that effectively operates as a high pass filter. The second derivative successfully eliminates the slowly-varying part, as can be seen in Fig. | cond-mat.mes-hall |
(< 1 % -6 % 132,133,134 126 A detailed analysis of the effective hole masses has been presented by Rodina et al. 127 Wurtzite AlN. The available experimental data on the band dispersion in wz-AlN is limited to the effective electron mass, which has been determined to be in the range of 0.29 to 0.45 m 0 . 123 The OEPx(cLDA)+G 0 W 0 values of m e = 0.32 m 0 and m ⊥ e = 0.33 m 0 fall within this range. Wurtzite InN. Experimentally derived effective electron masses in wz-InN scatter over a wide range (see Tab. | cond-mat.mtrl-sci |
From Eq. ( 14) it immediately follows that the interaction matrix element in Eq. ( 8) can be written as: I(k, q, q ′ , q+q ′ ) = I(0, q, q ′ , q+q ′ )f * (0, q, q ′ )f (k, q, q ′ ),(16) where all the k-dependence is contained in the function f (k, q, q ′ ) = e -i(γ k+q +γ k+q ′ -γ k+q+q ′ -γ k ) . (17) † m = 1 N φ k b † k e ik•rm ,(18) one can immediately see that the condensation (in the sense defined above) of the bosons into Bloch states corresponds to states with uniform phase winding along the basis directions of the triangular lattice in the Wannier basis, with the phase gradient given by the momentum k. This nature of the Abrikosov vortex lattice states has important consequences. First consequence, that can be seen immediately, is that the imaginary-time action, corresponding to longwavelength boson field phase fluctuations about a given Abrikosov state, will lack the usual (∇θ) 2 term, characteristic of superfluids, since all states with uniform phase gradients have the same energy. Instead, the action will have the form (after appropriate rescaling of the time and spatial coordinates): S ∼ dτ dr (∂ τ θ) 2 + (∇ 2 θ) 2 . | cond-mat.str-el |
While the origin of the unusual non-Fermi liquid resistivity seen under pressure in MnSi 23 (∆ρ ≈ AT 3 2 ) is not at present understood, the same temperature exponent is observed 25 in β-Mn, where it is expected to result from antiferromagnetic spin fluctuations. We believe it is worthwhile to investigate the possible link between the magnetic fluctuations and the non-Fermi liquid behavior in these materials. In this vein, it is interesting to note that powder neutron scattering down to 1.4 K 24 in β-Mn shows no signature of magnetic order. We are hopeful that this study might provide the motivation for single crystal neutron scattering to be carried out on β-Mn. In summary, we have used large-N theory for O(N ) vector spins and classical MC simulations to study the AF Heisenberg model on two three-dimensional cornershared triangle lattices, each site of which belongs to three equilateral triangles. The large-N studies suggested that the geometrical frustration present would lead to a partially ordered state on both lattices. | cond-mat.str-el |
Comment on "Self-Purification in Semiconductor Nanocrystals" In a recent Letter [1], Dalpian and Chelikowsky claimed that formation energies of Mn impurities in CdSe nanocrystals (NCs) increase as the size of the NC decreases, and argued that this size dependence leads to "self-purification" of small NCs. They presented densityfunctional-theory (DFT) calculations showing a strong size dependence for Mn impurity formation energies, and proposed a general explanation. In this Comment we show that several different DFT codes, pseudopotentials, and exchange-correlation functionals give a markedly different result: We find no such size dependence. More generally, we argue that formation energies are not relevant to substitutional doping in most colloidally grown NCs. We performed DFT calculations of the formation energies of Mn impurities in passivated CdSe NCs of different sizes, using methods similar to Ref. [1]. | cond-mat.mtrl-sci |
Though the insulating state is reminiscent of a Mott insulator, it is actually far from being one. The two 'Hubbard bands' are just the two spin bands shifted away from the Fermi level, and broadened to an extent that they just resemble non- interacting Gaussian density of states. In fact, the largefield spectral function may be deduced from the large field asymptote of equation 13. At large field h ∼ t * , the effective field becomes comparable to or larger than the largest scale in the problem, i.e U in strong coupling. This is because, for h ≃ t * , the magnetization saturates, m → 1, hence the effective field becomes at least h + U/2. The real part of the second-order self energy contributes 1/U at frequencies ω ≃ U , because Σ (2) (ω) ∼ 1/ω at large frequencies. | cond-mat.str-el |
During the nonadiabatic relaxation to the triplet state, which does not exclusively occur at the crossing point but also in its proximity where the two PESs do not coincide, the surplus of potential energy is reallocated onto all degrees of freedom. Since this allotment is performed along the nonadiabatic coupling vector (see Sec. II), not all the potential energy stored in the triplet-singlet gap E ts is transferred onto the translational degrees of freedom, which are principally responsible for surmounting the barriers. Rather, also the vibrational, rotational and lateral center-of-mass degrees of freedom, which do not promote the dissociation as strongly or might even hinder it, gain a certain portion of kinetic energy, which is then missing in the perpendicular translational motion. Consequently, these simulations are not directly comparable to the triplet MD, in which the complete amount E + ∆E ts is assigned to the translational degree of freedom perpendicular to the surface. The fact that there are two different mechanisms active at low incident energies depending on whether the minimal or the maximal coupling is employed becomes evident when inspecting the velocities that the triplet molecules exhibit in the reflection channel. | cond-mat.mtrl-sci |
Dilute moments also help support the sizeable coherence lengths (≃ 100 nm at 10 mK) observed in these materials. 9,10 Indeed, the presence of quantum interference effects in (Ga,Mn)As has been clearly demonstrated by measurements of universal conductance fluctuations and Aharonov-Bohm effects in (Ga,Mn)As nanodevices. 9,11 However, due to the aforementioned experimental subtleties conflicting conclusions have been reached 10,12,14 on the magnitude and even on the sign of quantum corrections to the conductivity. In this paper we report on a theoretical study which we expect to be helpful in achieving a more complete un-derstanding. Unlike earlier theoretical work 15,16 which addressed quantum interference in ferromagnets, we focus our study on a four-band model which is directly relevant to the valence bands of (Ga,Mn)As. We demonstrate that the quantum interference contribution to MR in robustly ferromagnetic (Ga,Mn)As is negative. | cond-mat.mtrl-sci |
2: The critical current density vs the thickness of the free (F1) layer. The circles are the experimental result of Chen et for 70×140 nm 2 junctions [12]. The solid line corresponds to Ic/S (see Eq. (11).) The dotted line and dashed-dotted line correspond to I 0 c /S and I p c /S, respectively. The dashed line corresponds to Ic/S in the limit of λt → 0. the component proportional to α pump . | cond-mat.mes-hall |
( 8) exactly agrees with magnetoelectronic circuit theory [4], but is now derived from spin pumping via the Onsager relations. We will in the following apply our theory to a disordered GaAs|(Ga,Mn)As|GaAs system, and investigate the magnetization dynamics in the ferromagnetic (Ga,Mn)As layer induced by an unpolarized charge current. We assume that the free-energy density depends on m as: F[m] = K c1 m 2 x m 2 y + m 2 x m 2 z + m 2 y m 2 z + K u m 2 z . (9) Here, K c1 is the lowest-order cubic anisotropy constant [12]. We assume the system is grown on a GaAs substrate that induces compressive strain in the (Ga,Mn)As layer. Strain breaks the cubic symmetry. | cond-mat.mes-hall |
Due to the translational invariance along the y-axis, the momentum k y is conserved. Hence, the common factor e ikyy is omitted above. By matching the wavefunctions at the interface x = 0 and L, we obtain the transmission coefficient t. We consider the situation that the barrier region is sufficiently narrow so that we can take the limit of U → ∞ and L → 0 while keeping Z ≡ U L = const. Here, we omit the expression of t because it is rather complicated but we note that it contains the barrier parameter Z only in the form of cos Z and sin Z. Consequently, the transmission probability and hence the conductance are π periodic with respect to Z. In the presence of Z, the spin direction of wavefunction rotates through the barrier region, similar to the spin transister. [7] Thus, with increasing Z, the connectivity of the wavefunction changes, which crucially influences the conductance.We parametrize k x +m x = k F cos φ, k y +m y = k F sin φ. | cond-mat.mes-hall |
When the distance to the contact is larger than the distance of the highest energy peak, the energy decreases monotonously as the dislocations approach the island center. Finally the defects stop at a finite distance from each other. We used the PFC model to make a similar calculation. We used mismatches of ±7.8% and pinning potential depth 0.5 times the amplitude of the triangular phase. The initial state is the upper half of a disk standing on the simulation cell edge where the pinning potential is effective. For compressive mismatch, the island does not form dislocations but adopts a concave shape with reduced size of the island-wetting film contact area. | cond-mat.mtrl-sci |
Minimum energy paths (MEP) associated with the creation and migration of kinks have been determined using the Nudged Elastic Band (NEB) technique. 47 Investigations of the formation and migration of kinks in bodycentered-cubic or L1 2 materials have already been done with a similar approach. 48,49,50 Both improved tangent and climbing images algorithms 51,52 have been employed in this work. Generally, we found that migration and formation mechanisms were simple enough to be described with a small number of images. Nevertheless, we used as many as 30 images in EDIP simulations, and 9 in first principles calculations. silicon atoms. | cond-mat.mtrl-sci |
1, we have valence and conduction bands with E g for the insulator phase. E g is monotonically getting smaller when the lattice constant a is being reduced by external pressure. If we have a first-order MIT, E g suddenly jumps from some positive value to the negative value (band overlap) at the critical lattice constant a = a c . This occurs if the total energy for the metal phase becomes lower than that of the insulator. Then we see electron pocket in valence band and hole pocket in conduction band, which are specified by the Fermi energy E F . Our analysis below shows that the first-order MIT is inevitably occur because of the nature of the long-range Coulomb interaction. | cond-mat.mtrl-sci |
If instead the optimization is performed for two adjacent matrices at a time, the resulting (quasi-variational) procedure is equivalent to White's original formulation of DMRG [5,10,11,12,13]. The MPS based formulation of this strategy has proven to be very enlightening and fruitful, in particular also in conjunction with concepts from quantum information theory [5]. In general, such an approach works for both bosonic and fermionic systems. However, to be efficient the method needs a local Hilbert space with finite and small dimension, limiting its applicability to cases where the local Hilbert space is finite dimensional a priori (e.g. fermions or hard-core bosons) or effectively reduced to a finite dimension, e.g. by interactions. | cond-mat.str-el |
(1.32) is Q = (π, 0). The variational wave function breaks the reflection symmetry of the lattice and, in finite systems, its energy can be lowered by projecting the state onto a subspace of definite symmetry. The results for the spin-spin correlations are shown in Fig. 1.4. By decreasing the value of J 2 /J 1 , we find clear evidence of a first-order phase transition, in agreement with previous calculations using different approaches [17,19]. For 0.5 < J 2 /J 1 < 0.65, the best variational wave function has no magnetic order (∆ AF = 0 and no Jastrow factor) and the BCS Hamiltonian has Energies per site for a 6 × 6 lattice. | cond-mat.str-el |
1 Introduction It is at the heart of the understanding of the high-temperature superconducting (SC) phase to infer what kind of state competes with the SC one in the possible quantum phase transition between the two. As in the two-dimensional (2D) copper oxides one of the candidates of such a state is the stripe phase, it is natural to expect that in the quasi-2D copper oxide coupled-ladder system also a kind of charge ordered state would compete with the predicted there SC state [4]. Indeed in Sr 14-x Ca x Cu 24 O 41 , which has the SC state for x = 13.6 and under pressure larger than 3 GPa [5], a charge density wave (CDW) phase with period λ = 3 and λ = 5 was observed when the Calcium concentration was tuned to x = 11 (which corresponds to the hole concentration n h = 6/5 per copper) and x = 0 (where n h = 4/3), respectively [1,2]. Although the charge transfer model for coupled ladders proposed in Ref. [3] may explain the onset of such an odd-period CDW state, it is the t-J-like model which is a natural Hamiltonian for the doped spin ladders [1,2]. However, then this experimental observation is a challenge for the theory since such a novel CDW state with the odd period has not been predicted by the standard t-J model for the ladders [6]. | cond-mat.str-el |
From the calculated MCPs the momentum space spin density has been reconstructed [Fig. 6(b)]. The calculated real space magnetization density map [Fig. 4(c)] has been obtained by a Fourier synthesis of the calculated magnetic structure factors. The main features of the respective density maps coincide well, although some differences are evident: the dip in the momentum space density around p z =0 is not pronounced well in the calculated map and has a shape, which is rotated by 90 • with respect to the observed map. This possibly results from strong hybridization effects between the Co3d and O2p orbitals. | cond-mat.mtrl-sci |
The deviations from the Drude conductivity may be read out from Fig. 1,i C δσ= α' -k+Q, β' -k+Q, k,α k,β k, β β'' k'', α'' -k''+Q, k,α β' -k+Q, α' -k+Q, δσ = e 2 2π dk v x α,β (k)v x β ′ ,α ′ (-k)G R α (k)G R α ′ (-k)G A β (k)G A β ′ (-k) dQ C β,β ′ α ′ ,α (k, Q)(1) where we set ≡ 1, α, β, ... label band eigenstates of the ferromagnet, v is the carrier velocity operator, ) is the advanced (retarded) Green's function in the first Born approximation, and Q is the "center of mass" momentum of the Cooperon. Q ranges approximately from the inverse phase coherence length, 1/l φ , to the inverse mean free path, 1/l. Following standard practice, we have kept Q in the Cooperon propagator only in Eq. ( 1), setting Q = 0 elsewhere in the integrand. Additionally, we ignore the contribution from non-backscattering processes, 26,27 which are unimportant in the diffusive regime. | cond-mat.mtrl-sci |
We thank T. Dahm and M. Takigawa for valuable discussions. Two of the authors (Y.N. and K.I.) thank K. Kitagawa and Y. Maeno for their experimental support. This work was supported by a Grant-in-Aid for the 21st Century COE gCenter for Diversity and Universality in Physicsh from the Ministry of Education, Culture, Sports, Science and Technology (MEXT) of Japan, by Grants-in-Aid for Scientific Research from the Japan Society for the Promotion of Science (JSPS), and by Grantsin-Aid for Scientific Research in Priority Area gSkutteruditeh (Nos. 16037208 and 15072206). | cond-mat.str-el |
It is remarkable that the gate dependence of the critical current on the Kondo ridges can be qualitatively described. That is in particular, the existence of a singlet, doublet (0/π) transition on ridge B at nearly the right value of gate voltage when renormalized with the charging energy U (arrows), and the absence of a (0/π) transition for ridge C corresponding to higher values of T K . One point of disagreement between theory and experiment concerns the amplitude of the critical current in the doublet state (π junction) region (around V BG = 3.65V). I c is theoretically found to be reduced by only a factor of two compared to its value in the singlet region whereas basically no trace of superconductivity could be detected experimentally. It is known however, that the π-phase current computed from the approximate FRG is too large compared to numerical renormalization group (NRG) data, which are known to be more accurate but only available at the center of the Kondo ridge 19 . Finally, let us emphasize the importance of the asymmetry of the transmission of the electrodes which tends to reduce considerably the supercurrent. | cond-mat.mes-hall |
The equation of motion derived was found to be essentially the same as that obtained by Berger long ago, 10,13 indicating his deep physical insights, but the effects of the spin-transfer torque and reflection force (momentum transfer) were obtained without phenomenological assumptions and ambiguities for the first time. Based on the obtained equation of motion, the wall motion under steady current was studied. It was found that in the adiabatic limit, where the reflection force can be neglected, and if in the absence of spin relaxation, there is a threshold current determined by the hard-axis magnetic anisotropy energy, K ⊥ . Thus the wall is intrinsically pinned by the internal degree of freedom, φ 0 . At large current, however, the wall gets depinned and its velocity becomes proportional to spin current (spin polarization of the current flow), j s , as is required from the angular momentum conservation. Numerical simulation was performed based on an equation of motion of each local spin by including the spin-transfer torque term in the adiabatic limit. | cond-mat.mes-hall |
We have computed the eigenenergies ε k in the range ε 0 ≤ ε k ≤ |ε 0 | for different QW numbers N (up to N = 20 ) at several sets of L and N L . It was found that for all those nanostructures the choice L = 500 and N L = 40000 enables one to satisfactory simulate the quasicontinuum states in the pointed energy interval. In this Section, we examine in detail the nanostructure composed of six QWs, see Fig. 3 (a). The plots of the excited state energies ε k (k=2 -7) on the interwell distance b (actually, on the barrier width) are presented in Fig. 3 (b). | cond-mat.mes-hall |
In addition, some signatures of pockets remain even without long-range order. The fluctuations are strong enough in this case to redistribute the spectral weight evenly between the real and the ghost FS. Our situation is intermediate in the sense that we have only quasi long range order, but the fluctuations have infinite correlation length. Our conclusion is that this situation is rather similar to the ordered case, i. e. the FS still has small pockets. The results of the Ref. [8] indicate that we have considered the strongest fluctuations possible consistent with small FS topology. | cond-mat.str-el |
Therefore, convergence is absolutely non-trivial. It turned out that convergence was best achieved with the Hartree-Fock approach, due to the fact that the band gaps are usually very large. Indeed, the band structure and corresponding densities of states display band gaps of ∼ 6 eV (Hartree-Fock), ∼ 0.5 eV (B3LYP), and ∼ 0.1 eV (LDA) for this initial structure. Note that this gap corresponds to a random initial structure and is very different from the gap of the final structure; but it is necessary to converge a calculation for this random initial geometry, and for all the geometries subsequently generated, until the end of the simulated annealing and quench. For comparison, calculations with the hybrid functional B3LYP were found to be much more difficult to converge, and a large mixing ratio was required: 90%, in combination with the Anderson mixing scheme; 35% was sufficient in the case of Hartree-Fock (the mixing ratio is the ratio with which the previous Fock operator is added to the new one, in order to achieve convergence). This leads to many iterations and thus a large CPU time. | cond-mat.mtrl-sci |
In addition to the unoccupied resonance, STM images and spatial maps of dI/dV acquired at different voltages evidence confinement of the resonance within the chain. In Fig. 7a spatially resolved dI/dV data acquired along the long symmetry axis of the chain at the indicated voltages are presented. Starting from a voltage which corresponds approximately to the resonance energy the dI/dV signal is piled up at the end of the chain whose position is indicated by a dashed line in Fig. 7a). In Figs. | cond-mat.mtrl-sci |
CONCLUSION The main results, obtained in this work, can be summarized as follows: i) In QD-chain with tunneling coupling in strong electron-photon coupling regime the space propagation of Rabi oscillations (Rabi waves) takes place. For propagation of Rabi waves the wave vector of the photon mode must have a nonzero component along the chain. Characteristics of the Rabi waves depend strongly on relations between parameter of electron-photon coupling, frequency deviation and transparency factors of tunneling barriers for both of levels. ii) Traveling Rabi waves are quantum states of the QDchain dressed by radiation. The qualitative distinction of this states from states of single dressed atom 2 is the space-time modulation of dressing parameter by traveling wave law. Traveling Rabi waves can be interpreted as entangled states of e-h pair and photons. | cond-mat.mes-hall |
In the following, primary fields are denoted by Vj , while the V j 's denote electron operators and are in general descendent fields. The ν = 2/5 ground state is given by 4 Ψ 2/5 = A N/2 i=1 V 1 (z i ) N/2 i=1 V 2 (z i )O bg ,(8) where A denotes antisymmetrization over the electron coordiates {z i }. In other words, the ground state is not just a single conformal block as in the Laughlin case, but rather an antisymmetrized sum over correlators involving two different representations of the electron. In analogy to the composite fermion picture, these two representations can be thought of as corresponding to composite fermions in the first and second CF Landau level, respectively. Since our formulation is entirely within the lowest Landau level, what distinguishes the two operators V 1 and V 2 is the presence of the derivative, which is naturally interpreted as giving the particle an extra "orbital" spin. This is in turn reflected in a shift in the relation between flux and particle number when transcribing the wave function eq. | cond-mat.mes-hall |
We begin our discussion with the standard 1 2 -filled band, for which the mechanism of the SP transition is well understood 57,64,65 . Fig. 7(a) shows in valence bond (VB) representation the generation of a spin triplet from the standard 1 2 -filled band. Since the two phases of bond alternation are isoenergetic, the two free spins can separate, and the true wavefunction is dominated by VB diagrams as in Fig. 7(b), where the phase of the bond alternation in between the two unpaired spins (spin solitons) is opposite to that in the ground state. With increasing temperature and increasing number of spin excitations there occur many such regions with reversed bond alternations, and overlaps between regions with different phases of bond alternations leads ultimately to the uniform state. | cond-mat.str-el |
For delocalized states, there are skew scattering Hall contributions due to impurity potentials. This has been treated in Refs. 33 and 35. There is also an intrinsic Luttinger contribution which can be written, in the language of Chazalviel 30 , and as rederived to first order in spin orbit coupling by Arsenault and Movaghar 25 . We write the transverse conductivity as σ i xy = e 2 Ω 1 e α - ∂f (ε α ) ∂ε α µ B σ α z ∆g zz α ,(17)where σ α z = n α↑ -n α↓ n α↑ + n α↓ ,(18) e = -|e| for electrons and e = |e| for holes. Finally, the zz component of the effective g-shift tensor is ∆g zz α = β α 4m 2 c 2 (∇V (r) × p) z β 1 ε β -ε α β |L z | α . | cond-mat.mtrl-sci |
This paper addresses this question using a toy-model built from square plaquettes; focusing mainly on the simplest case of spinless fermions, we explore the possibilities for realizing spontaneous currents. The main prior study of orbital currents in spinless models is Nersesyan's ladder model 19,20 , in which a map to spinfull chains was introduced that we adopt in Sec. IV B. Quite recently, spinless models were motivated by the possible realization in cold dilute atoms 21 . The choice of square plaquettes is a choice motivated both by convenience of calculation and real material geometries. As we will see, a square plaquette has spontaneous currents as one of its natural degrees of freedom which is what we desire to investigate: possibility of spontaneous currents in the zero-order ground state. | cond-mat.str-el |
(13)Here, g R (E) = [E1-H-v] -1 is the free-particle Green's function and Σ R (E) = ΣR (E) + Σ imp,R (E) is a retarded component of the self-energy. Detailed form of ΣR (E) is given in the Appendix. Whereas, the term of Σ imp,R (E) depends on diagonal components of it's own Green's function, as indicated by Eq. ( 8). Thus, we should solve the above matrix equation self-consistently. With the obtained G R and it's Hermitian conjugate G A , the Keldysh components of the Green's function and the self-energy become G K (E) = G R (E)Σ K (E)G A (E)(14)and Σ K (E) = ΣK (E) + Σ imp,K (E),(15) respectively. | cond-mat.mes-hall |
Results and discussion Tequila was injected at a rate of 6.26 x 10 -3 ml) per pulse from high to low pressure (4.5 psi to a few torrs) in an argon flow at 280 °C, prior to flash evaporation. Tequila basically consists of water and ethanol, both molecules will be dissociated through the reactor zone. Because of the bonding energies of C-C, C-H, C-O and H-OH, the ethanol molecule will be dissociated to supply the carbon atoms with hybrid bond (sp 3 ) , whereas water provides an excess of hydrogen to form other allotropes. The reactor temperature activates the silicon 3s and 3p orbitals to form the primitive partial covalent bond with the carbon hybrid orbital sp 3 . As the temperature increases, the surface reactions rates and the mobility of the adsorbed species are enhanced, thus increasing the growth rate [11]. Additionally, the surface roughness of the silicon substrate favors the nucleation process [10]. | cond-mat.mtrl-sci |
( 12) no longer holds. Level crossings are characterized by ∆S z = 2, and the saturation field H c2 is given by Eq. ( 11) with k = 2. This is exactly what would be expected where a spin nematic state is selected by quantum fluctuations in applied magnetic field 3 . However these are finite size results, and must be approached with a little caution. The critical fields associated with one-and two-magnon excitations show quite different finite size scaling properties as a function of φ, as illustrated in Fig. | cond-mat.str-el |
Finally, we have found that the β-fergusonite structure gives a phonon spectrum that can explain reasonably well our experimental results. According to grouptheoretical considerations the β-fergusonite has 18 Raman active modes at the Γ point: 8 A g + 10 B g . The frequencies and mode assignment of the different phonon calculated for this structure at 40 GPa are given in Table III. According with the calculations there are always two B g modes very close in frequency to each other, but lattice-dynamics calculations apparently tend to underestimate the frequency splitting between B g modes in fergusonite structures [19,20]. However, the calculated small splitting between B g modes could explain why in the experiments we have only found sixteen modes, since some of the B g modes could be degenerated within the accuracy of the experiments. The qualitative agreement between the calculated and measured phonon frequencies and pressure coefficients is reasonably good. | cond-mat.mtrl-sci |
Table 1 shows that the resonance frequencies of other SWNTs can be below the one expected from a doubly clamped beam. This may result from the additional mass of contamination adsorbed on the tube [1,2]. This may also be the consequence of slack, which occurs when the tube is longer than the distance between the electrodes [26]. To further investigate the effect of slack, we have introduced slack in a non-reversible way by pulling down the tube with the SFM cantilever. Figure 3(b) shows that f res can be divided by two for a slack below 1%. The slack s is defined as (L 0 -L)/L with L 0 being the tube length and L the separation between the clamping points. | cond-mat.mes-hall |
Rather, it is extrinsic, of tunnelling magnetoresistive (TMR) origin. Since the report of the large MR eect and high magnetic transition temperature, a number of experimental studies like NMR 13 , XES 14 , Hall measurements 15 , magnetic measurements 16 have been carried out to characterize various properties of this material. There have been also a number of theoretical studies involving both rst-principles calculations 17 -20 as well as model calculations 21 -26 . The unusually high ferromagnetic transition temperature in Sr 2 FeMoO 6 and related material like Sr 2 FeReO 6 was rationalized 17 ,23 in terms of a kinetic energy driven mechanism which produces a negative spin polarization at otherwise nonmagnetic site like Mo or Re. Following this idea, a double-exchange like two sublattice model was introduced and studied by dier-ent groups 21 ,22,24,25,26 . While most of the studies 21 ,24,25 were restricted only to ferromagnetic phase, some of the studies 26 ,22 were extended to other competing magnetic phases too. | cond-mat.mtrl-sci |
It shows three estimations of the local spectral function (many body density of states) for a theoretical model (described more fully below) of a Mott insulator: one obtained by an ED method and two obtained by maximum entropy (MaxEnt) 12 analytic continuation of imaginary time QMC data. While the qualitative structure of the three estimations appear consistent, there are significant differences of detail, including a factor of two in the size of the gap which makes it difficult to compare the theoretical results to data. In this paper we present a critical examination of different methods of determining the spectral function of a Mott insulator, and apply the results to the question of the gap value, spectral function and optical conductivity in the paramagnetic and antiferromagnetic phases of the two dimensional square lattice Hubbard model. We study MaxEnt analytic continuation of the Green's function and of the self-energy, and compare the results to ED calculations and to direct thermodynamic evaluations of the gap. We argue that continuation of the self energy provides the best method of minimizing the broadening effect of MaxEnt procedure. The self energy is also needed for computation of other response functions, for example the optical conductivity. | cond-mat.str-el |
point), the remaining static distortion increases as the Ca concentration is lowered, and as the fraction of sample that remains unmagnetized at 0.4T, increases. To investigate the correlations between local structure and magnetization, we plot the data in a new way 32 which simplifies the discussion. By combining σ 2 J-T /polaron (T) (Fig. 7) with the M(T) data (Fig. 1) we plot σ 2 J-T /polaron (T) vs M M 0 for the four samples in Fig. 8. | cond-mat.str-el |
In a previous study of Mg x C 60 , a stoichiometric compound 13 was reported for x = 4. In the present work we improved the synthesis and conclude from a series of samples with varying Mg content that the homogeneous phase lies in a range of Mg concentrations between x = 5 and x = 5.5. Mg 5 C 60 is the only example of an alkaline earth fulleride polymer. Previous studies of alkaline earth fullerides focused mostly on Ca x C 60 and Ba x C 60 (Refs. 14,15) superconductors which are not polymers. An early study of Mg doped fulleride films reported an insulating behavior for all Mg concentrations 16 in disagreement with our present results. | cond-mat.str-el |
The conventional view point is that the CEF states are composed by linear combination of Hund's rule ground state with J = 5/2. In this case, the highest rank of multipoles in this manifold is 2J = 5, which is too small to distinguish between the O h and T h symmetries. Namely, the Stevens operator of sixth rank O t 6 = O 2 6 -O 6 6 , which makes the difference, 11 has zero matrix elements with J = 5/2. On the other hand, recent experimental results suggest mixing of dipole and octupole degrees of freedom. 12,13 We analyze the wave functions in the CEF states taking higher order hybridization processes. It is found that the closeness of the J = 7/2 excited state above the J = 5/2 ground state tends to compensate the small ratio of hybridization over excitation energy of 4 f 6 intermediate states. | cond-mat.str-el |
This interest arose due to the various interesting ground states observed in this family of Ce-compounds. An unusually sharp phase transition near 27 K in the magnetic susceptibility of CeRu 2 Al 10 [10] has been attributed to a spin-dimer formation [14,15]. The resistivity of CeRu 2 Al 10 exhibits a sharp drop near 27 K resembling an insulator-metal transition [10][11][12]. A very similar phase transition, near 29 K, has been observed in CeOs 2 Al 10 [11,12], but in this compound the susceptibility (along aaxis) exhibits a broad maximum near 45 K in contrast to a sharp drop at the phase transition (27 K) in CeRu 2 Al 10 [10]. The broad maximum in the susceptibility and its strong anisotropic behavior in the paramagnetic state reveal the presence of strong hybridization between 4fand conduction electrons as well as strong single ion anisotropy arising from the crystal field potential. Another difference between the two systems appears in the resistivity of the ordered state; the resistivity of CeOs 2 Al 10 displays a thermal activation-type temperature dependence below 15 K while the resistivity of CeRu 2 Al 10 exhibits metallic behaviour below the phase transition down to 2 K. Furthermore, the observed phase transitions in these two compounds have some resemblance to that of the hidden order transition observed in URu 2 Si 2 with very small ordered state moment at 17 K [16]. | cond-mat.str-el |
( 8), it directly appears in the stationary state approached in the relaxation term (16). In a similar manner, we obtain Γ J = - J(x, t) -J s n(x, t) τ J . (17) For the simple system-bath coupling (15), which disregards backscattering processes, the relaxation times τ s and τ J turn out to be equal. In the following, we shall allow for different relaxation times τ J = τ s , which results when allowing for more general models involving backscattering processes [38], i.e., when elastic disorder is present. Before discussing the resulting spin torque, let us briefly comment on the validity regime of our description. Both the derivation of the relaxation terms, see Eqs. | cond-mat.mes-hall |
The epitaxial stabilization of perovskite SrIrO 3 phase enabled us to investigate the W-controlled changes in the electronic structures of the iridates systematically. We measured ab plane reflectance spectra of Sr 2 IrO 4 and Sr 3 Ir 2 O 7 single crystals at room temperature. The corresponding σ(ω) were obtained using the Kramers-Kronig transformation. For the SrIrO 3 thin films, we measured reflectance and transmittance spectra in the energy region between 0.15 and 6 eV. By solving the Fresnel equations numerically [17], we obtained σ(ω) for SrIrO 3 . Between 0.01 and 0.09 eV, we used far-infrared ellipsometry technique to obtain σ(ω) directly [18]. | cond-mat.str-el |
IV. The unexpected failure of the Hubbard-I approximation in the framework of NEGF 40,45 is analyzed in Sec. V. In section VI we extend the calculations beyond linear response by applying a generalized master equation formalism 46 which works in a basis of many-particle states and takes into account higher-order tunneling processes. In Sec. VII the failure of the mean-field Green function method for finite bias is demonstrated. Finally, we conclude on our findings in Sec. | cond-mat.mes-hall |
7 reported observation of thermally assisted depinning of a narrow domain wall under a current. Thermally-assisted current-driven domain wall motion has also been studied theoretically 8,9 . The present paper addresses current-induced magnetization noise in non-uniformly magnetized ferromagnets. The spatial variation of the magnetization direction gives rise to increased magnetization noise; by a fluctuating spin-transfer torque, electric current noise causes fluctuations of the magnetic order parameter. We take into account both thermal current noise and shot noise, and show that the resulting magnetization noise is well represented by introducing fictitious stochastic magnetic fields. By the fluctuation-dissipation theorem (FDT), the thermal stochastic field is related to the dissipation of energy, or damping, of the magnetization. | cond-mat.mes-hall |
In the same manner, Γ J follows from such a calculation. Acknowledgments We thank Gerrit Bauer, Rembert Duine, Mathias Kläui, and Gen Tatara for useful discussions. This work was supported by the SFB TR 12 of the DFG. APPENDIX A: RELAXATION KERNELS In this Appendix, we sketch the derivation of the relaxation kernels entering the equations of motion for the spin density and spin current density of the itinerant electrons. We use the abbreviations K s (x) ≡ s(x) and K J (x) ≡ J(x), and start from the master equation for the reduced density operator under the Markov approximation for our time-dependent system Hamiltonian H S (t) = H 0 + H ex (t). This master equation has been derived in detail in Ref. | cond-mat.mes-hall |
In this paper, we have investigated the three-orbital Hubbard model on the pyrochlore lattice in order to study the heavy fermion behaviors of LiV 2 O 4 . To study which type of degrees of freedom plays an important role in low-energy dynamics of this model, we have employed an approach of real-space renormalization group type. In the first stage of coarse graining, block variables are defined as follows for each primitive unit cell of pyrochlore lattice, i.e., a tetrahedron composed of four vanadium atoms. First we numerically diagonalized the three-orbital Hubbard model and calculated the ground state and lowenergy excited states in this unit for the cases of electron numbers from n d =4 to 7. The case of n d =6 corresponds to the average density in LiV 2 O 4 (d 1.5 per vanadium atom), and other cases describe charge excitations. One important result is that these low-energy states can be represented very precisely by a simple picture of molecular orbitals. | cond-mat.str-el |
While RPA correlation energy varies by only a fraction of a mRy (top panel), the lower panel shows a significant variation of several mRy for the exact-exchange energy. As discussed in Sec. III B, slow convergence of exact-exchange energy with respect to supercell size-or the density of the BZ sampling for an extended system-can result if, once the integrable divergence is eliminated by the Gygi-Baldereschi procedure, 18 the residual q = 0 term is not estimated correctly. This is shown in Figure 4 where the exact exchange energy of Be 2 molecule is shown as a of the inverse supercell volume. A large error, proportional to the inverse supercell volume, is present when the residual q = 0 term is simply neglected, while a much better convergence with system size is obtained when it is estimated according to the recipe described in Sec. III B. | cond-mat.mtrl-sci |
The 5d orbitals are spatially more extended than the 3d and 4d orbitals, so their W (U) values should be larger (smaller) than those of 3d and 4d TMOs. And the electron correlation should play much less important role. Therefore, many 5d TMOs have metallic ground states that can be described by band theory [5,6]. However, some 5d TMOs, such as Sr 2 IrO 4 , Sr 3 Ir 2 O 7 , and Ba 2 NaOsO 6 , have insulating ground states [7][8][9]. There have been some recent reports that correlation effects could be important for the 5d insulating TMO [9,10]. Then, we can raise a couple of important questions. | cond-mat.str-el |
This may constitute a limitation of the use of distributions in a localizing regime. However, the first and second moments are readily computed. The effective shear modulus µ is read from (26). With a ≡ f /π < 1/2, we have: M (1) µ µ = 1 -2a 1 + (m -1)a = 1 -(1 + m)(f /π) 1/2 + O(f ),(34a (ε PS ) = 1 -2a (1 -f )[1 + (m -1)a] , M (2) (ε PS ) = (m + 1)a f [1 + (m -1)a] ,(34b)S (1) (ε PS ) = (1 -2a)[(1 + f )a -f ] (1 -f )[1 + (m -1)a] , S (1) (σ PS ) = (1 + f )a -f (1 -f ) √ 1 -2a ,(34c)S (1) (ε m ) = m a(1 -2a) [1 + (m -1)a] √ 1 -f , S (1) (σ m ) = a (1 -f )(1 -2a) ,(34d)S (1) (ε SS ) = ∞, S (1) (σ SS ) = 0. (34e) The mean strain in the pore stems from M (2) (ε PS ) = [1 -(1 -f )M (1) (ε PS )]/f . Moreover, S(1) (σ PS ) = (µ/ µ)S (1) (ε PS ). | cond-mat.mtrl-sci |
We note that the slope of the confining potential is small when K J. It is nearly deconfined. Although the open string excitation is an energetically costly quasi-eigenstate, the flux-tube excitation of the closed string is less expensive. In the pyrochlore lattice, the shortest closed string is the pyrochlore hexagon the energy of which is √ 3K. In general, the energy of the closed string is l 2 K. Furthermore, we remark that the mechanism of the quantum confinement by the transverse field is non-trivial. As it is in the one-dimensional ferromagnetic Ising model, spinons are deconfined in the classical case and remain deconfined when the transverse field is turned on. | cond-mat.str-el |
III. RESULTS In our simulation of the electron interferometer we used r 0 = 1000nm for the radius of ring and for the electron we chose an effective mass m * = 0.067m 0 and wave number K = 0.1nm -1 . In addition to this we focus on B = 0 from here on since this is expected to produce the maximal spin interference between the two arms. We solved Eq. ( 5) numerically using the SVE approximation, and in order to check its validity we did the same calculation including the second order derivatives. The comparison between the two methods is shown in Fig. | cond-mat.mes-hall |
Whereas previous studies on such systems were dealing with very thin dots thus found in a nearly single-domain state [28], the thickness of our dots is in the range 50 -200 nm inducing fluxclosure states around a DW or vortex. These dots are perfectly suited for our needs because owing to the natural spread of shape occurring in self-assembly we can study the length of DWs as a function of the dot aspect ratio, by a statistical investigation of an assembly of dots over the same sample. For each dot we measured the experimental DW length, and computed the expected DW length predicted by the simple Van den Berg geometrical construction. This construction is relevant for vanishing thickness and infinite dimensions [29], and equals the dot asymmetry used in the simulations so that a direct comparison with the data of FIG. 2 is possible. FIG. | cond-mat.mtrl-sci |
As a strong check of the whole reported analysis, this comparison clearly demonstrates, by invalidating the presence of localized paramagnetic spins in BiBaCoO, that these dual electronic states are a source of an enhanced thermopower. Further experimental investigations are now needed in order to understand why these coexisting states are not observed in BiBaCoO in contrast to BiCaCoO. III. CONCLUSION To conclude, we have investigated the low temperature magnetic field dependence of the resistivity in the thermoelectric misfit cobalt oxide BiCaCoO from 60 K down to 3 K. The scaling of the negative magnetoresistance demonstrates a spin dependent transport mechanism due to a strong Hund's coupling. The inferred microscopic description implies dual electronic states which explain the coexistence between localized and itinerant electrons both contributing to the thermopower. By shedding a new light on the electronic states leading to a high thermopower, this result likely provides a new potential way to optimize the thermoelectric properties. | cond-mat.str-el |
The upright triangles depict the normalized 'average' magnetization M + Bχ collected on YbRu2Si2 [7]. The kink (shown by the arrow) in all the data is clearly seen in the transition region y ≥ 1. The solid curve represents yM * N (y) with parameters c1 and c2 adjusted for magnetic susceptibility of CeRu2Si2 at B = 0.94 mT. SCALING BEHAVIOR OF THE MAGNETORESISTANCEBy definition, MR is given by ρ mr (B, T ) = ρ(B, T ) -ρ(0, T ) ρ(0, T ) ,(9) We apply Eq. ( 9) to study MR of strongly correlated electron liquid versus temperature T as a function of magnetic field B. The resistivity ρ(B, T ) is ρ(B, T ) = ρ 0 + ∆ρ(B, T ) + ∆ρ L (B, T ), (10) where ρ 0 is a residual resistance, ∆ρ = c 1 AT 2 , c 1 is a constant, A is a coefficient determining the temperature dependence of the resistivity ρ = ρ 0 + AT 2 . | cond-mat.str-el |
Very recently 27 , the problem has been studied in detail in terms of a full numerical solution of spinfermion model and as well as in terms of reduced, classical magnetic model. These studies predict that when the competing magnetic phases are taken into account, the electron doped model systems beyond a certain doping prefers to have antiferromagnetic(AFM) arrangement of Fe spins compared to ferromagnetic(FM) arrangement of the undoped system. The predicted antiferromagnetic phase in electron-doped system is kinetic-energy driven rather than super-exchange driven, as is the case for example in Sr 2 FeWO 6 28 , which is an insulating anti- ferromagnet with Néel temperature of ≈ 20 K. The superexchange driven antiferromagnetic phase is necessarily insulating while the kinetic energy driven AFM phase may not be so. The prediction of such an antiferromagnetic phase of dierent origin is therefore of signicance. While the kinetic energy driven antiferromagnetic phases have been suggested in hole-doped rare-earth manganites (eg.the CE phase at half-doping 29 ), to the best of our knowledge, till date no reports of such analogous phases in double perovskites exist, thereby, opening up the possibility of experimental exploration in this front. However the afore-mentioned model calculations were carried out in two dimension and with single band, which was justied by the assumption that the dominant nearestneighbor B-B interactions are operative between orbitals of same symmetry and within a given plane. | cond-mat.mtrl-sci |
We have estimated the orbital (M orb ) and spin (M spin ) magnetic moments of the Ru 4d states using the XMCD sum rules as follows: M orb = -2 ∆A M3 + ∆A M2 3(A M3 + A M2 ) (10 -N 4d ),(1) M spin + 7M T = - ∆A M3 -2∆A M2 A M3 + A M2 (10 -N 4d ),(2) where M orb , M spin and the magnetic-dipole moment M T are given in units of µ B /atom, N 4d is the 4d electron occupation number which is assumed to be 4, ∆A M3 and ∆A M2 are the energy integrals of the 3p minimum. Since our measurements were made on polycrystalline samples, the angle average would result in a vanishing magnetic-dipole term 35 and therefore we have ignored M T in estimating the spin magnetic moment using Eq. ( 2). Fig. 4, we compare the XMCD intensities of Ru 3p 3/2 (∼ 461.5 eV) and O 1s (∼ 529 eV) with the magnetization measured at 2 T and the orbital and spin magnetic moments of Ru 4d estimated from the Ru 3p XMCD spectra. In the paramagnetic phase (x ≤ 0.3), however, no orbital magnetic moment is estimated from the Ru 3p XMCD spectra since no XMCD signals are observed in the Ru 3p within experimental error in as shown in Fig. | cond-mat.str-el |
RESULTS A. Ellipsometry and interband excitations In Fig. 2(a), we plot ε l = ε l 1 + ıε l 2 (l=a,c) from 0.75 to 5.8 eV. For convenience, the real part σ 1 = (ω/4π) ε 2 of the optical conductivity is displayed in Fig. 2(b). We find a striking anisotropy. In particular, there is only one strong peak at 5.6 eV in σ c 1 , while a multipeak structure is present in σ a 1 (peaks at ∼2, 3.5, 4.5, 4.9, and 5.5 eV). | cond-mat.str-el |
The paper is organized as following. In Sec. II, the macroscopic features of the magneto-gyrotropic effect, e.g., the possibility to generate a photocurrent in various experimental geometries and its behavior upon variation of the radiation polarization, are described in the frame of the phenomenological theory. In Sections III and IV, we give a short account of the experimental technique, present the experimental results on the photocurrents and discuss them in view of the theoretical background. Here, we also discuss applications of the MPGE, in particular, for the study of BIA and SIA responsible for the spin splitting of subbands in k-space. Preliminary results on the study of BIA and SIA are published in Ref. | cond-mat.mes-hall |
Raman scattering from HAFs can be understood in terms of the Loudon-Fleury (LF) processes 14 , in which two magnons are simultaneously created by light absorption and emission. In the limit of large on-site Coulomb correlations U , the Hamiltonian describing these processes can be obtained as a leading term of the expansion in t/(U -ω), where t is the nearest-neighbor (NN) hopping, and ω is of the order of photon frequencies 15 . The Raman intensity of HAFs on hypercubic lattices with unfrustrated NN exchange and collinear type of antiferromagnetic (AFM) order allows for a straightforward semiquantitative interpretation in terms of the LF processes. In fact, in real space, exchanging two NN spins of S = 1/2 leads to an excitation with energy Ω ∼ (z -1)J, where z is the coordination number and J is the AFM exchange energy. The reduction of Ω/J by -1 is a consequence of the exchange link between the NN sites and can be interpreted in terms of a twomagnon interactions in the final state. In momentum space, the linear spin-wave theory yields non-dispersive magnons along the the magnetic Brillouin zone (BZ) boundary, leading to a square-root divergence of the bare two-magnon density of states at Ω = zJ. | cond-mat.str-el |
After quenches to weak-coupling (U = 2, Fig. 9a), σ(t, t ′ ) undergoes a rapid initial relaxation, but it does not approach the thermal value within the accessible times. This behavior reflects the prethermalization that is observed in the momentum occupation. The conductivity at the corresponding effective temperature (T eff = 0.37) consists of a Drude peak at ω = 0 (Fig. 10), which is only slightly broadened due to temperature and interaction. Because a narrow Drude peak implies a slow decay of σ eq (s) with time difference, we cannot resolve the true width of the peak from data which are restricted to small times. | cond-mat.str-el |
The solid line through the data points is a fit obtained with parameters |h A | = 7.2 T and easy axis direction equal to 22 degrees from the film normal. These parameter values agree well with previous estimates based on equilibrium magnetization measurements [8,10]. Although the spectra in Fig. 2b are clearly associated with FMR, the sign change at low frequency is not consistent with Re χ ij (ω), which is positive definite. We have verified that the negative component is always present in the spectra and is not associated with errors in assigning the t=0 point in the time-domain data. The origin of negative component of the FT is made clearer by referring back to the time domain. | cond-mat.mtrl-sci |
17,18 Since the scalar order accompanies a slight lattice distortion, it can be probed by X-ray diffraction. 2 More detailed information should be obtained if 4 f form factors are probed by azimuthal scan in resonant X-ray scattering using the electric quadrupole (E2) channel. 19 Since the fourth-rank tensor relevant to E2 scattering is very complicated, we visualize the scalar order by deriving the simplest form factor that corresponds to a weighted average of the electron charge density. Namely, we utilize the integer (J = 4) value of the Pr 3+ configuration, and introduce a fictitious "wave function":ψ Γα (Ω) = Ω|Γ, α ,(10) where Ω represents the solid angle specified by (θ, φ) such that dΩ = sin θdθdφ. Then ψ Γα (Ω) can be derived in terms of spherical harmonics Y 4m (Ω) with use of eqs. ( 1) to ( 9). | cond-mat.str-el |
The Keldysh Green's function, which is the basic quantity of the KB approach, of the open wire is studied in Section III A showing different timedependent regimes relevant to the subsequent analysis. In Section III B and III C we calculate the TD current and dipole moment respectively. We find that the 2B and GW results are in excellent agreement at all times and can differ substantially from the HF results. We also perform the Fourier analysis of the transient oscillations and reveal the underlying out-of-equilibrium electronic structure of the open wire. 62 The dynamically screened interaction of the GW approximation is investigated in Section III D with emphasis on the time-scales of retardation effects. Section III E is devoted to the study of the TD rearrangement of the density in the two-dimensional leads after the switch-on of an external bias. | cond-mat.mes-hall |
Thus ∆ Q = 0, which means there isn't rectification in the linear response regime. This corresponds to the one obtained in [33] for the thermoelectric transport in a chain of quantum dots with selfconsistent reservoirs. On the other hand, if the ratio τ RC (ω)/τ RL (ω) is a constant for any frequency ω, by using QR = 0, one can easily find ∆ Q = 0 and the rectification is absent even in the nonlinear response regime. Therefore, even in the nonlinear regime, the thermal rectification is absent in the symmetric three-terminal junctions with τ RC (ω) = τ RL (ω). In a more general situation where the ratio τ RC (ω)/τ RL (ω) for any frequency is not a constant, the fact that τ RC (ω)/τ RL (ω) varies with ω means there are different asymmetries for transmission of phonons in different frequencies. One can certainly expect that ∆ Q is not zero in the nonlinear response regime and it is determined by the dependence of the ratio τ RC (ω)/τ RL (ω) on ω. | cond-mat.mes-hall |
Recently, topological insulators suitable for room temperature applications are also predicted 4 . The QSH state has the helical edge states, namely, having two counter-propagating edge states for the two opposite spin polarizations. The helical edge states are stable against time-reversal conserving perturbations, since backscattering processes need to connect the upper and lower edges of the sample. The probability of backscattering is exponentially suppressed as the sample width is increased. Recent experiment 5 provides evidences of the QSHE in HgTe/CdTe quantum well (QW) structures, as predicted theoretically 6 . The de-coherence effect in QSH samples is also investigated. | cond-mat.mes-hall |
Later Read and Green developed the analogy between the MR state and a p-wave superconductor and showed that the Ising-type topological charge was related to a Majorana fermion localized on the quantized vortices. This analogy was elaborated further in papers by Ivanov 39 , and Stern et al. 40 . Referring back to our previous discussion about fractional statistics, it should be clear that any conclusions about the non-Abelian fractional statistics based on the monodromies of the wave functions depend on an implicit assumption that the Berry phases vanish. Just as in the case of the hierarchical states, this conjecture has not been rigorously proven, but strong arguments in favor have recently been given by Read 34 . Additional numerical evidence has been given by Tserkovnyak and Simon 41 and by Baraban et al. | cond-mat.mes-hall |
To proceed, we diagonalize the quadratic part of the Hamiltonian H 2 by a Bogoliubov transformation to a set of magnon quasiparticles a k = u k c k + v k c † -k (11) a † k = u k c † k + v k c -k , where c ( †) k are bosons, and the coherence coefficients u k = A k + E k 2E k (12) v k = - B k |B k | A k -E k 2E k . satisfy u 2 k -v 2 k = 1. The Hamitonian H 2 in terms of the Bogoliubov quasiparticles reads 10) -( 12) of Ref. 17 H 2 = k E k c † k c k ,(13) and the dispersion is given by E k = A 2 k -B 2 k = (1 -ν k )(1 + 2ν k ) . (14) The magnon dispersion E k is depicted in Fig. 2. | cond-mat.str-el |
The basic reasoning for this is presented in the previous subsection: since the value of V is in certain sense small, the first few terms of the perturbation series with respect to V can make sense. More detailed discussion about the small parameters of the theory are presented in the next sections; let us first present the general properties of the diagrams under consideration. The rules of diagram construction are quite similar to the usual Matsubara diagram technique. The only difference from the standard perturbation scheme is that the interaction operator V is not purely of the 4-th order form f * f * f f , and therefore vertices in the diagrams are not necessarily four-leg, but may formally have any even number of legs. For the choice (13), these vertices are essentially γ (n) . They are connected with the lines being the dual Green's functions. | cond-mat.str-el |
B. Encounters touching the lead opening The calculation of K(t) shown above was based on a semiclassical calculation of the covariance of the reflection coefficients off the two contacts. This method to calculate K(t) is technically simplest because one only needs to consider encounters that reside in the interior of the sample. The encounters that appear in the calculation of K(t) (a) and K(t) (c) can touch the contacts, however, if a different expression is used to calculate K(t). The encounters that appear in the calculation of K(t) (b) never touch the contact because they are part of a periodic trajectory. In order to illustrate a semiclassical theory for K(t) with encounters that touch the lead opening, we calculate K(t) from the covariance of reflection and transmission coefficients, using g(ε) = N 1 -tr S 11 (ε)S 11 (ε) † , g(ε ′ ) = tr S 12 (ε ′ )S 12 (ε ′ ) † (88) | cond-mat.mes-hall |
The latter can be successfully applied [5] to derive results (3)-( 5). This approach was recently extended further to find T 3 1 and T 1 (1 -T 1 )T 2 (1 -T 2 ) exactly [9]. Here, we have calculated all moments with i n i 4 by using tricks of partial integrations to reduce all moments to forms of Selberg's integral. We will report on that in more detail elsewhere [10]. By this method we are able to calculate the so-called skewness (third cumulant) of the conductance, which we represent in the following compact form: g 3 ≡ (g -g ) 3 = var(g) 4[(1 -2 β ) 2 -(N 1 -N 2 ) 2 ] β(N -3 + 2 β )(N -1 + 2 β )(N -1 + 6 β ) . (8) It is worth noting that the skewness vanishes for symmetric cavities (N 1 = N 2 ) at β = 2. | cond-mat.mes-hall |
DFT 10,11,12 is a practical approach to many-body physics which recognizes the impossibility of achieving exact results and seeks practical solutions with adequate accuracy. Following a familiar line of argument 10,11,12 which we do not reproduce here, many-body exchangecorrelation effects can be taken into account in the graphene many-body problem with the same formal justifications and the same types of approximation schemes as in standard non-relativistic DFT 10,11,12 . The end result in the case of present interest is that ground state charge densities and energies are determined by solving a timeindependent Kohn-Sham-Dirac equation for a sublatticepseudospin spinor Φ λ (r) = (ϕ (A) λ (r), ϕ (B) λ (r)) T , [vσ • p + I σ V KS (r)] Φ λ (r) = ε λ Φ λ (r) . (1) Here v ∼ 10 6 m/s is the bare Fermi velocity, p = -i ∇ r , σ is a 2D vector constructed with the 2 × 2 Pauli matrices σ 1 and σ 2 acting in pseudospin space, I σ is the 2 × 2 identity matrix in pseudospin space, and V KS (r) = V ext (r) + ∆V H (r) + V xc (r) is the Kohn-Sham (KS) potential, which is a functional of the ground-state density n(r). The ground-state density is obtained as a sum over occupied Kohn-Sham-Dirac spinors Φ λ (r): n(r) = 4 λ(occ) Φ † λ (r)Φ λ (r) ≡ 4 λ(occ) [|ϕ (A) λ (r)| 2 + |ϕ (B) λ (r)| 2 ] ,(2) where the factor 4 is due to valley and spin degeneracies and {ϕ (σ) λ (r), σ = A, B} are the pseudospin (sublattice) components of the Kohn-Sham-Dirac spinor Φ λ (r). Equation ( 2) is a self-consistent closure relationship for the Kohn-Sham-Dirac equations ( 1), since the effective potential in Eq. | cond-mat.str-el |
These are the fundamental reasons to look for a feasible and controlled way of the valley-isospin manipulation in graphene. The practical consequences of eliminating fermion doubling effects are even more exciting: The spin-based quantum computing in graphene quantum dots [3], exploiting the superior spin coherence expected in carbon nanostructures, is one of the most recent examples. More generally, the potential of graphene for future electronics rests on the possibility to create devices that have no analogue in silicon-based electronics [4]. Briefly speaking, it seems that valley degree of freedom might be used to control an electronic device, in much the same way as the electron spin is used in spintronics [5] or quantum computing [6]. Therefore, a controllable way of occupying a single valley in graphene (in other words: producing a valley polarization) would be a key ingredient for the so-called "valleytronics". The author, Tworzydło, and Beenakker recently proposed an electrostatic method of valley polarization in a single-mode quantum point contact with zigzag edges [7]. | cond-mat.mes-hall |
Of all the compounds and phases discussed in this article wz-GaN is the best characterized experimentally. The good agreement between our quasiparticle band structures and those based on the parameter set recommended by VM'03 proves the quality of our OEPx(cLDA)+G 0 W 0 band structures. wz-AlN the effective electron masses recommended by VM'03 are the averages over several theoretical values; the recommended VB parameters are theoretical values by Kim et al. 135 derived from LDA calculations. These parameters yield a band structure, which is in good overall agreement with the OEPx(cLDA)+G 0 W 0 band structure (see Fig. 3a). | cond-mat.mtrl-sci |
To understand this, we provide a qualitative analysis of the coupling strengths of individual eigenstates to the leads. At first, all states with even m (e.g. Figs. 2.f, 2.g and 2.i) do not contribute to the transmission as they possess negative transversal y-parity, implying zero overlap with the propagating states in the leads, such that the zero field transmission is determined by the interplay of (n, m) states with odd m. Later we will see that the states with even m make one contribution to the magnetoconductance for non-zero field strength. In the following, we distinguish between confined states (CS) and leaking states (LS). A CS does not fill the whole area of the oval with a substantial probability amplitude but is located near the center x = 0 and decays rapidly in the direction towards the lead stubs (e.g. | cond-mat.mes-hall |
However, in the subsequent dynamics the O atom emerges again above the surface and the gained kinetic energy is gradually transferred via the lattice vibrations to the whole surface slab. After quenching of the atomic motion, the O atoms are stably adsorbed in hollow surface sites, separated by a distance of 5.9 Å, i.e. by more than two Al-Al distances. In a number of similar FPMD simulations, the same "hot-atom" dissociation mechanism has been observed to lead to spontaneous O 2 dissociation on Al(100), Ti(0001), Co(0001), Cr(110), TiN(001), and notably also on Si(001) [36]. We may thus consider it as a general mechanism for the initial reaction of dioxygen molecules with bare metal or semiconductor surfaces. Among the system investigated so far, only the reaction of dioxygen with Pt(111) resulted in stable adsorption of molecular O 2 without dissociation. | cond-mat.mtrl-sci |
End of preview. Expand
in Dataset Viewer.
README.md exists but content is empty.
Use the Edit dataset card button to edit it.
- Downloads last month
- 1