ISSP

Activity Report

2007

Contents

Pages

1

PrefaceResearch Highlights
Highlights of Joint ResearchInternational Conferences
and Workshops
ISSP Workshops
Subjects of Joint Research
Publications

2 - 3536 - 51
52 - 53
54 - 57

58 - 115

116 - 162

ISSP Activity Report 2007

Preface

This booklet constitutes the annual report of the activityof the Institute for Solid State Physics (ISSP) of the Universityof Tokyo for the fiscal year 2007. The ISSP was establishedin 1957 as a joint-use central research institution in the fieldof basic materials science, so that the year 2007 was the 50th
anniversary.
In the late 90’ s, ISSP made a big decision to leave the
Roppongi campus after its 40 years of operation there inorder to seek a fresh restart in the new Kashiwa campus.
In preparation of the restructuring and relocation, the ISSP
underwent its first major external evaluation by an international
review committee in 1995. The comments and suggestions givenby the review committee were highly valuable in envisioning
the future of the ISSP in Kashiwa. The construction of researchbuilding started in 1997 and the first group of laboratories began
research operation in Kashiwa in April 1999. The relocation of
the main body of the ISSP followed and the full operation in
Kashiwa started in April 2000.
In the fall of 2005, the ISSP had the second major external evaluation by an international review
committee. In the review, evaluation was given on the research accomplishments, especially thoseutilizing new facilities installed at the time of relocation. I believe it is fair to say that the scientificachievements during the first 5 years of research activity in Kashiwa were highly praised by thereviewers. Valuable comments and suggestions were also given on the future research directions and thebetterment of the ISSP management.
In November 2007, ISSP held the 50th anniversary and a commemorative symposium in the old
Roppongi campus, which has now been renovated as the National Art Center and the National Graduate
Institute for Policy Studies. For many old-timers, this was a kind of “home-coming” event. In the
symposium, the “past, present and future” of ISSP and materials science were discussed.
In addition to the various on-going research activities, there are a few near-future projects
ranging from the actual construction stage to the planning stage. The one currently in the final stageof construction is the long-pulse high magnetic field laboratory, which, when completed, will offera versatile platform for experiments in high magnetic field, and will constitute a comprehensive
International MegaGauss Science Laboratory for high magnetic field research together with the existing
short-pulse megagauss field installation. Another project vigorously promoted is the Advanced Coherent
Light Laboratory which is envisaged as a laser facility enabling experiments in the unexplored realm of
short wavelength, ultrafast and high peak power.
The ISSP is now taking a step forward to its second 50 years. It is a good occasion for us to reflect
on the mission, past accomplishments, and future strategy of the ISSP, so as to keep playing the leadingrole for the advancement of basic science in the relevant fields by the synergy of the in-house staffs andthe research communities.

June, 2008

Yasuhiro Iye
Director

Institute for Solid State Physics
The University of Tokyo

ISSP Activity Report 2007

Research Highlights

A Novel Ordered Phase in the
Frustrated 2D Spin System
SrCu2(BO3)2 under High Pressure
Takigawa, Y. Ueda, and Uwatoko Groups
A variety of exotic quantum phenomena has been discovered in SrCu2(BO3)2, a quasi two-dimensional frustrated spin
system with a dimer-singlet ground state (Fig. 1ab), mostnotably, the sequential magnetization plateaux at 1/8, 1/4,
and 1/3 of the saturation. These plateaux accompany superlattice of triplets that breaks the translational symmetry of
the crystal, as observed by NMR experiments [1].
Compared with the behavior in high magnetic fields,
little has been known about the properties under highpressure, although pressure would be a powerful tool toinvestigate quantum phase transitions expected for spin
systems with such a geometry. We have performed nuclear
magnetic resonance (NMR) experiments on 11B nuclei in asingle crystal of SrCu2(BO3)2 under pressure up to 2.4 GPa,
which revealed a magnetic phase transition resulting ina novel type of valence-bond-solid (VBS) order. A pistoncylinder type pressure cell was mounted on a doubleaxis goniometer, which enabled precise alignment of themagnetic field along arbitrary crystalline directions.
The NMR spectra at 2.4 GPa in the field of 7 T along
the c-axis are shown in Fig. 2 at various temperatures (T ).
A sharp single line is observed at high T consistent with the
crystal symmetry. The line, however, splits gradually below
30 K. The angle dependence of the spectrum indicates thatthe splitting is due to loss of C4 symmetry of the crystal,
which makes the two orthogonal dimer sublattices inequivalent (solid and dashed blue lines in Fig. 1a). With decreasing
temperature, each line further splits into one sharp andone broad lines below 3.6 K. The magnetic hyperfine shift(K ) of the sharp line shows an activated T-dependence
indicating an excitation gap, while the broad line shows
gapless behavior with nearly constant K (Fig. 3a). The clear

Fig. 1. The magnetic layer of SrCu2(BO3)2 viewed along (a) thec-direction and (b) the [110]-direction. The Cu2+ ions with spin 1/2
form orthogonal dimer sublattices as shown by the solid and thedashed blue lines. (c) Proposed VBS order of alternating magnetic(shaded) and non-magnetic (non-shaded) dimers.

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ISSP Activity Report 2007

Fig. 2. NMR spectra for H // c at various temperatures. A gradualline splitting below 30 K is followed by another sudden splitting ofeach line into one sharp and one broad lines below 3.6 K.

Fig. 3. (a) Temperature dependence of the magnetic hyperfine shift.
The black open squares show the data for unsplit lines at high T,
while the blue and red open squares represent the shifts for split
lines in the intermediate T-range. The filled symbols indicate the shift
in the ordered phase. For comparison, the data at ambient pressure
are shown by the green dots. (b) Field-temperature phase diagram
obtained from the NMR data.

singularity in both K(T ) (Fig. 3a) and the susceptibility
indicates a well defined magnetic phase transition.
The detailed investigation of the angle dependence ofthe spectra allowed us to conclude that each dimer sublattice develops a superstructure with alternation of magnetic
and non-magnetic dimers, resulting in a doubled unit cell(Fig. 1c). The magnetic dimers appear to restore a finite gap
at low magnetic fields below 2 T. Hence the ground state
at zero field would be a novel type of valence-bond-solidstate with staggered dimer correlation, which breaks thetranslational symmetry but does not show any spontaneousmagnetization. The transition temperature decreases withmagnetic field (Fig. 3b), suggesting a quantum phase transition near 18 T. Further experiments to clarify the nature ofthis totally unexpected ordered phase are in progress.
References[1] K. Kodama, M. Takigawa, M. Horvatic, C. Berthier, H. Kageyama,
Y. Ueda, S. Miyahara, F. Becca, and F. Mila, Science 298, 395(2002).
[2] T. Waki, K. Arai, M. Takigawa, Y. Saiga, Y. Uwatoko,
H. Kageyama, and Y. Ueda, J. Phys. Soc. Jpn. 76, 073710 (2007)(Slelected for JPSJ Editor’ s Choice).
Authors
T. Waki, K. Arai, M. Takigawa, J. Yamaura, Y. Saiga, Y. Uwatoko,
H. Kageyamaa, and Y. Ueda
aDepartment of Chemistry, Kyoto University

Gap Determination of the Heavy Fermion
Superconductor URu2Si2 by Specific
Heat Measurements in Magnetic Fields
Sakakibara Group
The tetragonal compound URu2Si2 has attracted muchinterest, primarily because it exhibits an exotic ordering at
To=17.5 K whose order parameter is still an issue of intense
debate and referred to as a “hidden order parameter” [1].
This hidden order coexists with superconductivity thatsets in below Tc=1.4 K [2]. Regarding the superconducting
properties of this system, the existence of nodes in the
superconducting gap has been reported in several experiments including the specific heat, NMR and NQR measurements. The pairing symmetry, however, has remained
unidentified. In order to examine the gap structure of thesuperconducting state, we have carried out field-orientation
dependent specific heat C(H) measurements on a single
crystal of URu2Si2 (Tc =1.3 K, 100 mg weight) [3].
Figure 1(a) shows C(H) of URu2Si2 for various orientations between the c- and a-axes. It is remarkable thatthe anisotropy of C(H) is weak at low fields, but becomespronounced when H exceeds about 1 T where the effect ofanisotropy in the upper critical field Hc2 becomes significant. Accordingly, the field dependent part of the specificheat CH as a function of H/Hc2 becomes strongly field-

Fig. 1. (a) Field variations of the specific heat of URu2Si2 obtained at
0.34 K for various field-orientations between c and a axes. (b) Fielddependent part of the specific heat for the two principal directions
plotted as a function of H/Hc2.

Fig. 2. Schematic gap structures for the superconducting state of
URu2Si2. (a) is the gap structure suggested by the CH measurements.
Two point nodes are located in the z direction. Group theoreticalconsiderations require an additional line node as shown in (b), which
is probably invisible due to the light effective mass of the quasiparticles.

orientation dependent at low fields as shown in Fig. 1 (b).
When H ‖ a, CH exhibits a steep increase at low field witha clear downward curvature, similar to the H1/2 behaviorof nodal superconductors due to the “Doppler effect” ofmoving quasiparticles (QPs) near the gap nodes [4]. As forH ‖ c, however, CH almost linearly increases at low fields,
which is the behavior expected for a full-gap superconductor. More generally, this means that there is no nodal QP
that is Doppler-effect active for this field orientation. These
two qualitatively different field dependences of CH can bereconciled with each other if point nodes are located along
the z directions as depicted in Fig. 2 (a). In such a case, thenodal QPs cannot contribute to the zero-energy density ofstates when H ‖ c, because the momentum of those QPs isperpendicular to the superfluid velocity (circulating supercurrent around the vortices), making the Doppler effectinactive.
As can be seen in Fig. 1 (b), CH for H ‖ c exhibits adistinct upturn above H ~ 0.4Hc2. In fact, a linear extrapolation of the low field data of CH (the dotted line in the figure)reaches the normal state value at H/Hc2 ~ 1.2. This behavior
is in contradiction with the case of ordinary (nodal as well
as full-gap) superconductors where the similar extrapolation
reaches the normal-state value at H/Hc2 < 1. This unusual
behavior of CH in URu2Si2 can be explained by a strong
Pauli paramagnetic depairing effect that suppresses Hc2 tothe value well below the orbital limiting field [3].
These observations lead us to conclude that the pairing
symmetry of the superconducting state in URu2Si2 is evenparity (singlet pairing) with point nodes in the z (polar)
direction (Fig. 2 (a)). Group theoretical considerations fora tetragonal crystal tell us that a possible even parity gap
function with linear point nodes is kz(kx+iky). This gap
function, however, possesses a line node on the equator(kz=0) (Fig. 2 (b)), whose presence was not confirmed in our
experiment. We believe that this discrepancy comes from
Fermi surface topology of the present system. It should benoticed that C(H) measurements mainly probe the quasiparticles in the heaviest band. If the line node in question is
located on a band with a light effective mass, it would not
be detected by the C(H) measurements. While we have toconsider realistic Fermi surfaces for conclusive evaluation of
the gap structure, linear point nodes in the polar direction
are quite plausible in URu2Si2.
References[1] H. Amitsuka et al., J. Mag. Mag. Mater. 310, 214 (2007).
[2] T.T.M. Palstra et al., Phys. Rev. Lett. 55, 2727 (1985).
[3] K. Yano et al., Phys. Rev. Lett. 100, 017004 (2008).
[4] G.E. Volovik, JETP Lett. 58, 469 (1993).

ISSP Activity Report 2007



Authors
K. Yano, T. Sakakibara, T. Tayama, M. Yokoyamaa, H. Amitsukab, Y.
Hommac, P. Miranovicd, M. Ichiokae, Y. Tsutsumie, and K. Machidae
aIbaraki University
bHokkaido University
cTohoku University
dUniversity of Montenegro
eOkayama University

Electroluminescence Quenching
Caused by a Spin-crossover Transition
Tajima GroupThe spin-crossover transition in metal-coordinate
system interests many researchers from the view pint ofvarious functionalities such as color change and photo
response (LIEST) [1]. In order to utilize them in an electric
device, thin-film fabrication is necessary. In this respect,
we have fabricated a thin film of [Fe(dpp)2](BF4)2 (dpp =2,6-di(pyrazol-1-yl)pyridine) by employing the spin-coating
technique. This compound exhibits a transition around260 K in a single crystal [2]. The fabricated films were
~30 nm in thickness with roughness less than 5 nm.
Temperature dependence of the magnetization of the films
exhibits a spin-crossover transition around 260 K. We also
observed changes associated with the spin-crossover transition in optical absorption spectra [3].
Very recently we fabricated organic light-emitting diodeusing [Fe(dpp)2](BF4)2 as the host compound and chlorophyll a as the guest. We found an electroluminescencequenching caused by the spin-crossover transition. The
phenomenon is reproducible for thermal cycles. This is thefirst observation of a switching phenomenon in an organic
thin film device accompanying the spin-crossover transition[4].
Figure 1 shows the device architecture and the structures of molecules used for device fabrication. The thickness of the emissive layer was approximately 50 nm. The

Fig. 1. The device architecture, molecular structures of [Fe(dpp)2]2+
and chlorophyll a, and the image of thin film of [Fe(dpp)2](BF4)2.

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ISSP Activity Report 2007

Fig. 2. Electroluminescence spectra of the ITO/chlorophyll
a:[Fe(dpp)2](BF4)2/Al device at different temperatures under theapplied voltage of 3.5 V. The inset shows the temperature dependence of the EL intensity at 689 nm.

roughness of the surface was less than 5 nm.
Figure 2 shows the EL spectra at different temperatures under the applied voltage of 3.5 V, and the inset
depicts the temperature dependence of the EL intensity
at 689 nm where the emission maximum was observed.
The EL emission from the fabricated ITO/chlorophyll a:[Fe(dpp)2](BF4)2/Al device under the applied voltage of 3 V
was clearly observed at 300 K, and the spectrum was thesame as that of the single-layer device with chlorophylla [5]. However, as the temperature decreased to 200 K,
the EL spectrum of chlorophyll a could not be observed;
heating to 300 K enabled the spectrum to be obtained
again. This ON–OFF switching of the EL emission is reproducible for the heating and cooling processes. As thetemperature decreases, the intensity of the EL emission
decreases, and around 260 K, the intensity becomes comparable with that of the background level (< 260 K).
Since the temperature of 260 K is almost the same as thespin transition temperature of [Fe(dpp)2](BF4)2 complex, it
is considered that the EL quenching at 260 K is relevant tothe spin transition. One possible origin is an energy transferfrom excited chlorophyll a to [Fe(dpp)2](BF4)2. As a resultof the spin transition, the energy level of [Fe(dpp)2](BF4)2
should change. If the energy level shifts from above
to below that of the excited chlorophyll a, the energy
transfer that leads to the nonradiative decay process canoccur. In order to examine this possibility, we measured
the photoluminescence (PL) spectra for the chlorophylla: [Fe(dpp) 2](BF 4) 2 film at 300 and 200 K. Although
the PL intensity observed at 200 K was weaker than thatat 300 K, the PL spectrum of chlorophyll a could beobtained at 200 K in contrast to the EL spectra vanishing.
This result rules out the possibility that the quenching of
the EL emission from chlorophyll a is caused by the energy
transfer, and suggests the disappearance of the excitedchlorophyll a in the EL device. We speculate that carrier
injection efficiency to chlorophyll a is drastically changed
as the spin transition of [Fe(dpp)2](BF4)2 occurs. In otherwords, in the low-temperature region where [Fe(dpp)2](BF4)2
is at a low-spin state, injected carriers pass through theinsulating layer via [Fe(dpp)2](BF4)2 without generating theexcited state of chlorophyll a. Detailed studies are underway.

References[1] For example, see "Spin Crossover in Transition Metal
Compounnds III", ed by P. Gütlich et al., (Springer, 2004).
[2] J. M. Holland et al., J. Chem. Soc., Dalton Trans. 2002, 548[3] M. Matsuda and H. Tajima, Chem. Lett. 36, 700 (2007).
[4] M. Matsuda et al., Chem. Lett. 37, 374 (2008).
[5] K. Shimatani et al., Chem. Lett. 34, 948 (2005).
Authors
M. Matsuda, H. Isozaki, and H. Tajima

Giant Nonlinear Conduction andElectric-Field-Induced Metastable
State in an Organic Conductor
β-(meso-DMBEDT-TTF)2PF6
Mori GroupIn the past few decades, nonlinear conductions inorganic conductors have been extensively studied inneutral-ionic transition complexes, Spin-Peierls materials,
Mott insulators, π - d systems, and others. Recently, as a
new type of nonlinear conductors, the charge ordered (CO)organic materials are attracting much attention because
they require low voltage to activate their nonlinear conductions. The low activation voltage enables us to effectively
exclude the problems of Joule heating and may help studies
on intrinsic nonlinear conduction. Moreover, the peculiar
thyristor-like I-V characteristics and DC-AC converter effectof θ - (BEDT-TTF)2CsM ’ (SCN)4 (M ’ =Zn, Co) once again
promoted the study of nonlinear conduction phenomena,
especially in CO materials [1]. To investigate the relationship between charge ordering and dielectric response andto study nonlinear conduction in CO materials, we carriedout measurements on dielectric constant and I-V characteristics in an organic CO material, β - (meso-DMBEDT-TTF)2PF6[2-5].
Figure 1 shows I-V curves obtained by the I-driven
4-probe method. Nonlinear conduction is observed ina wide temperature range below 100 K. Especially, thenegative differential resistance (NDR: dV/dI < 0) can be seen
below about 70 K. It should be noted that the voltage drop
pauses around 20 A/cm2 and then the small NDR appears
around 40 A/cm2. To clarify the singularity in its I-V curves,
we fit the results of the 4-probe measurement by thephenomenological equation:σ ( J,T ) = σ1 exp(-Δ/T)+σ2 Jn		(1)

Fig. 1. I-V curves obtained by the I-driven 4-probe method forβ-(meso-DMBEDT-TTF)2PF6. Solid lines are the fit to eq. (2).

Fig. 2. Apparent sample resistivity Esample/J as a function of Vcircuitfor β-(meso-DMBEDT-TTF)2PF6.

Fig. 3. Time-resolved sample voltage Vsample at 60.8 K for β-(mesoDMBEDT-TTF)2PF6.

The first term reflects the ordinary Arrhenius-type temperature dependence of conductivity. The second term means
the current-induced enhancement of conductivity. By introducing the threshold current density JT [dE/dJ(J=JT)=0] we
can rewrite this equation as follows:σ ( J,T )=σ0[1+( J / JT)n /(n-1)](2)where σ0= σ1exp(-Δ/T ). We fitted the experimental datawith the fitting parameters σ0, JT, and n in the range ofJ ≤ 10 A/cm2. From the fitting, we obtained JT = 2.5~4.7
A/cm2 and n = 1.8~2.0. This n value is larger than ordinary
nonlinear conductors with the n value of ~1.5. The solidlines in Fig. 1 show the fit to eq. (2). We can see from thefigure that the experimental data deviate upward from thefitting function in the high-current region. According to eq.
(2), the conductivity increases with the power of J, which
expects the linear behavior with the slope of n in the larger
J region. However, the slope varies with the J value in therange of J ≥ 10 A/cm2 and cannot be determined uniquely.
To compare with the results of the V-driven 2-probe
method, we plotted the apparent sample resistivity
Esample/J obtained from the 2-probe method as a functionof Vcircuit as shown in Fig. 2. At 55.5 K, for example, theresistivity shows a steep drop by a factor of 103 at thethreshold voltage of 4 V. Such a giant nonlinear conduction with a low threshold voltage is similar to that of
θ-(BEDT-TTF)2CsCo(SCN)4 and could be attributed to themelting of the CCO (checkerboard CO). It is also noted thatthe subsequent melting process occurs even after the giant
resistivity drop above 4V.

ISSP Activity Report 2007

To further investigate the peculiar two-stepped I-Vcharacteristics, we monitored the sample voltage under
the pulsed circuit voltage using an oscilloscope (Tektronix
TDS2012B). The results are shown in Fig. 3. If the circuit
voltage Vcircuit is small, the sample voltage Vsample keeps
a constant value near Vcircuit owing to its large resistivity(R~1×105 ohm, ρ~80 ohm cm). When Vcircuit reaches about
2 V, Vsample begins to drop suddenly with some delay after
the pulse application and the larger Vcircuit causes the faster
drop of Vsample. In the short time span just after the steep
voltage drop, we can see a terrace of low voltage (t = 1ms,
Vsample = 0.9 V at Vcircuit = 3.00 V, for example) before
dropping into the more conductive state. Such a non-monotonus response has never been observed in other materials
studied so far, and this plateau would be the reason for
its peculiar two-stepped NDR. Another noticeable pointis that the LR-CCO (longrange-CCO below 70 K) shows a
two-stepped drop with the metastable plateau, while the
SR-CCO (shortrange-CCO above 70 K) gives a small, monotonous drop.
In conclusion, we observed the formation of LR-CCO
as a steep increase in the frequency-dependent dielectric
constant at 70 K. In the transport measurement, (1) the
nonlinear I-V curve below 100 K, the peculiar negative
resistance around 70 K, and a giant resistance drop by afactor of up to 103 under 4 V at 55.5 K were observed,
and in the time-resolved measurement technique (2) atwo-stepped sample voltage drop with a transient plateau
occurred below 70 K. This indicates that the melting of
SR-CCO gives a small and monotonous voltage drop,
whereas the LR-CCO changes drastically to a temporary
plateau before dropping to the most conductive state.
Especially, this plateau below 70 K implies a new metastablestate that can be induced only by the electric field.
References[1] F. Sawano, I. Terasaki, H. Mori, T. Mori, M. Watanabe, N. Ikeda,
Y. Nogami, and Y. Noda, Nature 437, 522 (2005).
[2] S. Kimura, T. Maejima, H. Suzuki, R. Chiba, H. Mori,
T. Kawamoto, T. Mori, H. Moriyama, Y. Nishio, and K. Kajita, Chem.
Commun.2454 (2004).
[3] S. Kimura, H. Suzuki, T. Maejima, H. Mori, J. Yamaura,
T. Kakiuchi, H. Sawa, and H. Moriyama, J. Am. Chem. Soc. 128,
1456 (2006).
[4] H. Mori, J. Phys. Soc. Jpn. 75, 051003 (2006).
[5] M. Tanaka, K. Yamamoto, M. Uruichi, T. Yamamoto, K. Yakushi,
S. Kimura, and H. Mori, J. Phys. Soc. Jpn. 77, 024714 (2007).

Fig. 1. The photoimage of the fabricated HTS SQUID.

fields less than 5~mT, not suitable for the magnetization
reversal measurements. Here, we report on the operationof HTS micro SQUID with a thinner film and a smaller loop
than SSMs to enable measurements while sweeping themagnetic field produced by a 3D magnetic field coil system.
The hysteresis loop of micro ferrimagnetic crystal has
successfully been measured in any direction of magnetic
field up to 0.12 T between 30 and 70 K [1, 2].
A dc micro SQUID with grain boundary Josephson
junctions (GBJJs) was fabricated according to the standard
procedure. By using a pulsed laser ablation technique,epitaxial 92 nm thick YBa2Cu3O7-δ (YBCO) films were
deposited onto a [001]-oriented SrTiO3 (STO) bi-crystal
substrate (10 × 102 mm2) with misorientation angle of 24°.
The YBCO film deposition was followed by in-situ sputter
deposition of 70 nm-thick Au layer for electrical contact
and protection during the subsequent processes. The dc
micro SQUID was formed by using the photolithography
and the ion milling dry etching techniques to have a hole of
4 × 9 μm2 and two Josephson junctions with 2 μm in width(W) across the bi-crystal line (Fig. 1). The fabricated chip
was mounted on a copper holder with a thermometer(Lakeshore CERNOX 1030) and a heater to control thetemperature from 4.2 K to 300 K. The holder is supported
from the top of vacuum can (VC) by several rods of phenolic
resin (PPS) for moderate cooling and efficient temperature

Authors
S. Niizeki, F. Yoshikane, K. Kohno, K. Takahashi, H. Mori,
Y. Bandoa, T. Kawamotoa, and T. Moria
aTokyo Institute of Technology

High Temperature Superconductor
Micro SQUID Magnetometer for
Magnetization Measurementof a Micro Scale Magnet
Mori and Ishimoto Groups
So far there are several reports on the HTS SQUID
techniques, where the bare SQUID loop as well as the micro
SQUID system is used as the detection coil. The scanning
SQUID microscope (SSM) can observe stray magnetic field
images produced by small wires of ~20 μm in diameter.
However, these measurements were made only in fixedISSP Activity Report 2007

Fig. 2. Schematic diagram of the measurement probe unit with a 3D
magnetic field coil system.

control. The 3D magnetic coil system (B1, B2, and B3)placed outside the VC consists of three mutually orthogonalcoils. This allows accurate sweeping of the magnetic field in
the desired orientation (Fig. 2).
The present system was used to measure the magnetic
hysteresis of a micron-sized magnetic material. In the pointdipole approximation, a dipole moment μ (JT-1) produces aflux at the SQUID loop:
ϕ = BsAeff = 3μ μ0 Aeff sinθcosθ / 4π|r|3
Here, Bs is the magnetic flux density from the sample
placed at the center of the SQUID loop, Aeff=77.4 μm2,
μ 0 the free-space permeability constant, r the vector
joining the sample to the SQUID loop, and θ the angle
between r and the direction of the dipole moment. The
sample here is a molecular ferrimagnet: [Mn 2 (H 2 O) 2
(CH3COO)][W(CN)8]·2H2O with Tc = 40 K. Panels (a) and (b)in Fig. 3 represent Ic vs B2 at 35 and 50 K, respectively. At
35 K, the SQUID modulation period precipitously drops as
the field approaches 0 T, while at 50 K it is almost constantover the whole field range. Furthermore, as is seen in theinset of Fig. 3(a), the central peaks in the Fraunhofer likepattern at 35 K exhibit negative- and positive-field shifts inthe field-decreasing and -increasing process, respectively.
There is no such behavior at 50 K. This may imply a generation of remanent magnetization at 35 K. Second harmoniccomponents appear in the modulation in the range of -24 to-10 mT in Fig. 3(d) and (e). The period is likely reduced by afactor of two compared with a normal one. It leads to 1/2ϕ0
period wherever the second harmonic component appears.
In this way the magnetization curves are obtained within
an accuracy of less than 1/2ϕ0. There can be seen a clear
contrast between the curves above and below Tc in Fig.
3(c). The magnetization curve at 50 K is linear with respect
to the applied magnetic field. It is mainly from the strayfield since the paramagnetic component in the sample isexpected to be negligible. Thus, we can extract the sample
magnetization curve at 35 K (the saturated value of ~40 ϕ0around 30 mT) by subtracting the 50 K curve from 35 Kone.
In summary, we have developed an HTS micro SQUID
magnetometer combined with a 3D magnetic field coilsystem that is designed for investigating the magnetization
reversal of tiny magnetic materials. It can be operated in
the temperature range between 30 and 70 K (an improvement over LTS micro SQUIDs) and under the external field
up to at least 0.12 T. The anisotropic behavior of hysteresis
loop has successfully been measured for a micro crystalline molecular ferrimagnet. This is the first time that a
HTS micro SQUID under magnetic field up to above 0.1
T. The present sensitivity is an order of magnitude higher
than MPMS but far below the submicron Hall magnetometer operable in similar temperature range. If the distance
between the sample and the SQUID loop were reduced bya factor of ten in the present case, the sensitivity would be
expected to increase roughly thousand times as much as
now.
References[1] K. Takeda, H. Mori, A. Yamaguchi, H. Ishimoto, T. Nakamura,
S. Kuriki, T. Hozumi, and S. Ohkoshi, J. Applied Phys. 103, 07E911(2008).
[2] K. Takeda, H. Mori, A. Yamaguchi, H. Ishimoto, T. Nakamura,
S. Kuriki, T. Hozumi, and S. Ohkoshi, Rev. Sci. Instr. 79, 033909(2008).
Authors
K. Takeda, H. Mori, A. Yamaguchi, H. Ishimoto, T. Nakamuraa,
S. Kurikia, T. Hozumib, and S. Ohkoshib
aHokkaido University
bDepartment of Chemistry, the University of Tokyo

2D Frustrated Magnetism vs.
3D Spin Order in the Chalcogenide
Triangular Antiferromagnets
NiGa2S4, FeGa2S4, and Fe2Ga2S5
Fig. 3. Ic vs B2 plots at 35 K (a) and 50 K (b) with the sample
mounted on SQUID II. The black and gray lines show the fieldreducing and -increasing processes, respectively. The inset to (a)gives details around 0 T. The black-circle and gray-triangle markersin (c) show the magnetization curves of the sample deduced bycounting the number of the detected flux depending on the appliedmagnetic field at 35 and 50 K, respectively. At each temperature
the filled (open) markers denote the curves in the field-reducing(-increasing) processes. The behavior in the range of -24 and -10 mTin the field-reducing (d) and -increasing processes (e), respectively.

Nakatsuji Group
Geometrically frustrated magnets have attracted greatinterest for the possible exotic ground states and phasetransition phenomena. The two dimensional triangularlattice is one of the simplest forms of a geometricallyfrustrated lattice with a single magnetic ion in a unit cell.
NiGa2S4 is the first example of S = 1 two dimensional(2D) antiferromagnets on an exact triangular lattice [1].

ISSP Activity Report 2007



References[1] S. Nakatsuji, Y. Nambu, H. Tonomura, O. Sakai, S. Jonas,
C. Broholm, H. Tsunetsugu, Y. Qiu, and Y. Maeno, Science 309,
1697 (2005).
[2] H. Takeya et al., Phys. Rev. B 77, 054429 (2008).
[3] D. E. MacLaughlin, Y. Nambu, S. Nakatsuji, R. H. Heffner, L. Shu,
O. O. Bernal, and K. Ishida, submitted to Phys. Rev. Lett. (2008).
[4] S. Nakatsuji, H. Tonomura, K. Onuma, Y. Nambu, O. Sakai,
Y. Maeno, R. T. Macaluso, and Julia Y. Chan, Phys. Rev. Lett. 99,
157203 (2007).
Authors
S. Nakatsuji, Y. Nambu, Y. Maenoa, and J. Y. Chanba Kyoto University
b Louisiana State University

Fig. 1. (a) Four sublattices of the buckled honeycomb lattice formedby the Fe-S-Fe straight bonds in the bilayer of Fe2Ga2S5. (b), Scaling
in the T 2 dependence in the magnetic part of the specific heat CM
for NiGa2S4 (S = 1; ΘW = 80 K) and FeGa2S4 (S = 2; ΘW = 160 K) at 0
T in full logarithmic scale (c) ab-plane and c-axis components of thesusceptibility χ for FeGa2S4 and Fe2Ga2S5 under B = 0.1 T. Inset χ(T)and dχ /dT of Fe2Ga2S5. The vertical lines indicate the Néel point TN.

Despite an antiferromagnetic coupling with the Weisstemperature ΘW = 80 K, no antiferromagnetic or canonicalspin glass ordering has been observed down to 0.08 K[1, 2]. Instead, recent nuclear-quadrupole-resonance andmuon-spin relaxation measurements have revealed unconventional freezing phenomena that set in below Tf ~ 10 K
and have a wide temperature critical fluctuation regime
down to 2 K [2, 3]. Very interestingly, below Tf ~ 10 K, it
develops a short-range noncollinear order with 2D gaplesslinearly dispersive excitations characterized by a T2 dependence of the specific heat [1]. In order to explain the origin
and stability of the 2D coherent behavior in the frozen
spin-disordered state, various theoretical proposals havebeen made. Experimentally, it is highly important to clarify
the effect of spin size change, bilayering, and geometricalfrustration, and to see if any spin order can be realized in
the related compounds.
Thus, we have made the single crystal study of theisostructural single-layer triangular antiferromagnets,
NiGa2S4 and FeGa2S4, and the homologous bilayer triangular antiferromagnet Fe2Ga2S5 [4]. We have succeededin growing single crystals of NiGa2S4 for the first time andfound properties fully consistent with those for polycrystalline samples. Interestingly, the magnetic properties of
FeGa2S4 bear strong resemblances to those of NiGa2S4
despite the fact that Fe2+ has twice larger S = 2 spin thanS = 1 for Ni2+. Both compounds have basically Heisenberg
spins, and form frozen disordered state (see Figure 1). In
addition, the specific heat is insensitive to the field andshows a T 2 dependence, which scales to that of NiGa2S4(see Figure 1(b)). The similarities strongly suggest that the
2D coherent behavior in the frozen spin-disordered state
has the same underlying mechanism. In contrast, a clear
antiferromagnetic transition is observed for the bilayer
system Fe2Ga2S5, whose dominant antiferromagnetic bonds
most likely form an unfrustrated honeycomb lattice (seeFigure 1(a) and (c)). These results suggest that the geometrical frustration of the single-layer triangular lattices stabilizes the spin-disordered state, probably associated with 2D
antiferromagnetic ordering [4].

ISSP Activity Report 2007

Nonperturbative Approaches toCorrelated Transport throughQuantum Dots under a Bias Voltage
K. Ueda GroupOne of challenging problems in mesoscopic physics is
to understand many-body effects in a nonequilibrium situation. A central issue is the Kondo effect at quantum dots.
By applying a finite bias voltage, quantum dots enable us
to investigate the nonequilibrium Kondo effect. We have
employed the perturbative Keldysh Green function method,by which it has been settled whether the Kondo peak splits
or not [1]. Toward full understanding of electron correlation
effects at a steady state, nonperturbative treatments are also
required. In recent years we have addressed this issue from
two aspects.
A. Shot noise and nonequilibrium Kubo formula based on
Keldysh formalism for mesoscopic systems
As for the Keldysh formalism applied for mesoscopic
systems, we have examined an adiabatic switching-on of
the perturbation term. We have proven that the procedurecorresponds to taking the invariant part introduced by
Zubarev. As a consequence the density matrix becomes oneof MacLennan-Zubarev forms. This concept enables us toderive a general form of differential conductance, given by
the sum of the current-current correlation function and thenon-trivial current-charge correlation function. This general-

Fig. 1. The differential conductance under a magnetic field for
U/t=2, where t is the hopping amplitude in the leads, U the Coulomb~
energy at the dot, V the bias voltage and h the Zeeman energy. It
is well known that the split peak is located at the Zeeman splitting(twice of the Zeeman energy), and our result shows an excellent
agreement.

Importance of Nonlocality in the
Dynamical Exchange-Correlation
Kernel in the Time-Dependent
Density Functional Theory
Takada Group

Fig. 2. The expectation values of the double occupation operator
taken by the steady state wave function for U/t=2. Here we can see a
non-monotonic behavior for relatively small magnetic field.

ization may be called as the nonequilibrium Kubo formula.
It is well known that the current-current correlation
function gives noise power at zero frequency. The generalized Kubo formula leads us to define shot noise by the
non-trivial current-charge correlation function. This fact isconfirmed in case of the noninteracting Anderson model.
The nonequilibrium Kubo formula have thus provided
a new perspective on the shot-noise study [2]. With theresulting formula of shot noise, we are currently studying
the shot noise in the Kondo effect at quantum dots.
B. Time-dependent DMRG study on quantum dot under afinite bias voltage
Apart from Keldysh formalism, it seems that there are
only few numerical studies on quantum dot under a finite
bias voltage, due to a lack of reliable numerical techniques
for nonequilibrium problems. One possible numericalapproach is the adaptive time-dependent density matrix
renormalization group (TdDMRG) method. In this method,
the ground state wave function is calculated with the usual
DMRG method initially. Then the time evolution of the wave
function due to the slowly changing bias voltage between
the two leads can be calculated by using the TdDMRG
technique. Even though the system size is finite, the expectation value of the current operator shows steady likebehavior for a finite time interval, in which the system is
expected to resemble the real nonequilibrium steady state of
the infinitely long system.
Taking this approach, we have shown that thus calculated wave function in the time interval contains accurate
information on the properties of real steady state of theinfinite system in a wide range of bias voltage [3]. We havesuccessfully reproduced the splitting of the zero bias peakin the differential conductance, Fig. 1. Moreover, an anomalous behavior has been observed in the expectation valuesof the double occupation operator at the dot as a function ofbias voltage, Fig. 2. This method is applicable also to othernonequilibrium problems of mesoscopic systems.
References[1] T. Fujii and K. Ueda, Phys. Rev. B 68, 111310 (2003); T. Fujii andK. Ueda, J. Phys. Soc. Jpn. 74, 127 (2005).
[2] T. Fujii, J. Phys. Soc. Jpn. 76, 044709 (2007).
[3] S. Kirino, T. Fujii, J. Zhao, and K. Ueda, submitted to J. Phys.
Soc. Jpn. (2008).
Authors
T. Fujii, S. Kirino, J. Zhao, and K. Ueda

The time-dependent density functional theory (TDDFT)
is a powerful tool for studying excitations in atomic, molecular, and condensed-matter systems. In the linear-response
regime, the key quantity of TDDFT is the dynamical
exchange and correlation (xc) kernel fxc(r,r’ ;ω), defined as
the Fourier transform of fxc(r,r’ ;t-t’ ) which is given by thefunctional derivative of the Kohn-Sham xc potential Vxc(r,t)
with respect to the electron density n(r’ ,t’ ). In actual implementation of the theory, we need to resort to some approximation to this quantity. In the absence of its detailed information, the adiabatic local density approximation (ALDA)
has been adopted in virtually all calculations so far.
Some years ago, we derived a rigorous formula forobtaining the experimentally observed stopping power,–dE/dx, of solids for slow ions in the framework of TDDFT[1], but it turns out that the formula provides a nonzero
value for the friction coefficient (or –dE/dx divided by ν thevelocity of incident ions in the limit of ν → 0) in the homogeneous electron gas with the limit of zero electron density(or in the vacuum), if we adopt ALDA for fxc(r,r’ ;ω). This
contradictory result urges us to develop a better approximation scheme for this quantity beyond the level of ALDA.
In pursuit of such a scheme, we have reformulated the
problem of the stopping power in the framework of thetime-dependent current density functional theory (TDCDFT)[2] to find a new interesting rigorous relation between
fxc(r,r’ ;ω) and the tensorial xc kernel, the key quantity

Fig. 1. Comparison between experiment on the friction coefficientof aluminum for slow ions (characterized by the atomic number)
obtained at the incident velocity ν = 0.5 atomic units (the solid
squares) and theory on the same quantity of the homogeneouselectron gas with the electron density equivalent to the valenceelectron density in aluminum in the limit of zero ion velocity(ν → 0). The open circles, the open triangles, and the open squares
correspond, respectively, to the results in our new approximation
scheme (LDA to CDFT), ALDA to fxc(r,r’ ;ω), and LDA to the static
Kohn-Sham scheme.

ISSP Activity Report 2007

in TDCDFT. By invoking the local density approximation
to the tensorial xc kernel and exploiting this relation, we
have constructed a new approximation form for fxc(r,r’ ;ω).
The resulting form adds a nonlocal nature to ALDA andthe nonlocality is just enough to make the result for thefriction coefficient of the electron gas free from the abovementioned contradiction in the vacuum limit.
In order to check the accuracy of this new form forfxc(r,r’ ;ω), we have recalculated the friction coefficient of
the homogeneous electron gas at various electron densities
and compared the obtained results with the experimental
ones. Note that our results for the homogeneous electron
gas can be directly compared with the experimental ones in
simple metals, in view of the fact that at low velocities theenergy loss of ions in solids is due mainly to the stopping
power of valence electrons.
In Fig. 1, an example of such comparisons is shown at
the electron density as specified by rs=2.2 (rs: the conventional nondimensional density parameter) corresponding
to aluminum with incident ions penetrating from its (111)
surface. In the figure, we have also plotted the results in
ALDA as well as those in LDA to the static Kohn-Shamscheme which amounts to the calculation without including
any dynamical xc effects.
As can be seen in Fig. 1, experiment exhibits an oscillatory behavior in the friction coefficient with increasing
the atomic number of incident ions. This behavior, which is
attributed to the shell structure of valence electrons aroundthe ion, is reproduced well in all the calculation schemes,
but the best quantitative agreement with experiment isobtained by our new scheme (namely, LDA to CDFT) forfxc(r,r’ ;ω).
We e x p e c t t h a t o u r n e w a p p r o x i m a t i o n f o r m
fxc(r,r’ ;ω) awaits various applications in a wide range ofproblems concerning dynamical processes in electronicsystems.
References[1] V. U. Nazarov, J. M. Pitarke, C. S. Kim, and Y. Takada, Phys.
Rev. B 71, 121106(R) (2005).
[2] V. U. Nazarov, J. M. Pitarke, Y. Takada, G. Vignale, and Y.-C.
Chang, Phys. Rev. B 76, 205103 (2007).
Authors
V. U. Nazarov,a J. M. Pitarke,b Y. Takada, G. Vignalec, and Y.-C.
ChangaaAcademia Sinica, Taiwan,
bEuskal Herriko Unibertsitatea, Spain
cUniversity of Missouri, USA.

Fig. 1. Magnetization of the S =1 Heisenberg antiferromagnetic chain(Haldane chain) as a function of temperature, for various values ofapplied magnetic field above the critical field. Each curve shows aminimum of magnetization at a finite temperature. Such a minimum
is absent in the magnetization of a classical chain (shown in red).
The temperature at the minimum increases, as the field increases.

field. At zero temperature, the system undergoes a quantum
phase transition at the critical field where the gap closes.
Above the critical field lies a gapless phase. In two andhigher dimensions, the gapless phase at zero temperature has a long-range order of transverse magnetization.
However, in one dimension, quantum fluctuations are
so strong that the long-range order is absent even in thegapless phase. Instead, the gapless phase can be describedas Tomonaga-Luttinger liquid, which is the universal theory
for many gapless systems in one dimension. In fact, theS = 1/2 Heisenberg antiferromagnetic chain, which isgapless already at zero field, can be also described as theTomonaga-Luttinger liquid.
This implies that, once the gap is closed by the appliedfield, a one-dimensional spin gap system would become
essentially identical to intrinsically gapless systems, inthe low-energy limit at zero temperature. However, if we
consider finite-temperature properties, “remnants” of spin
gap appear in various ways. We examined numericallythe temperature dependence of the magnetization in the

Universal Magnetization Minimum in
One-Dimensional Spin Gap Systems
under Magnetic Field
Oshikawa Group
Spin gap systems have been vigorously studied bothexperimentally and theoretically over decades. Spin gap
systems generally lack usual magnetic order even at zerotemperature, thanks to quantum fluctuation. Thus they
provide an ideal playground to study effects of quantum
fluctuations. Various spin gap systems have been foundexperimentally, including one-dimensional ones such as
Haldane gap systems and two-leg spin ladders.
The spin gap can be closed by an applied magnetic

10

ISSP Activity Report 2007

Fig. 2. The crossover diagram on the field-temperature plane. In one
dimension, phase transition is absent at finite temperature. However,there can be a crossover between different regimes. In the present
case, as the temperature is increased at a constant field, the systemcrossovers into the “quantum critical regime” governed by the
non-relativistic, parabolic dispersion of magnons.

S = 1 Haldane chain under magnetic field, using Quantum
Monte Carlo simulation. Above the critical field, we find aminimum in the magnetization as a function of the temperature as in Figure 1. Such a minimum does not appear in themagnetization of the intrinsically gapless S = 1/2 Heisenberg
chain. In fact, the magnetization minimum can be understood as a signature of crossover between the “quantum
critical” regime governed by the nonrelativistic quantum
critical point and the Tomonaga-Luttinger liquid regime, as
in Figure 2. This is different from the similar magnetization
minimum in three dimensional systems, which corresponds
to Bose-Einsten condensation of magnons.
We have revealed that the magnetization minimum isa universal phenomenon for field-induced gapless phases
slightly above the critical field. In particular, the temperature Tm at the minimum, the applied field H, and the criticalfield Hc obey the relation Tm = x0 (H – Hc) when H is close
to Hc, where x0=0.76238…. is a universal constant.
Reference[1] Y. Maeda, C. Hotta, and M. Oshikawa, Phys. Rev. Lett. 99,
057205 (2007).
Authors
Y. Maedaa, C. Hottab, and M. Oshikawa
aUniversity of British Columbia
bKyoto Sangyo University

Pt/water Interface －towards a Full
Understanding of the Electrochemistry－
Sugino GroupIn our daily life, we utilize many electrochemical
reactions, such as, electrolysis of water, fuel-cell reactions,corrosion, and plating/etching. These phenomena would be
used to ultimately solve many problems related to environment, clean energy, and bio-membrane when the electrode/solution interface could be completely controlled in the
nano-meter scale. However, the interfacial phenomena areso complicated and controlling the interface is not easy,

Fig. 1. The electrode dynamics at Pt(111). ESM-FPMD simulation on
the Volmer step (H3O+ + e − → H(ad) + H2O) of the hydrogen evolution reaction. The hydronium ion H3O+ at the contact releases ahydrogen atom toward the surface accompanying the discharging. To
accelerate the simulation, negative bias of − 0.7 V vs. the standardhydrogen electrode (SHE) was applied to the interface.

Fig. 2. The layer distribution function of the water molecules (oxygensite). With increasing negative bias, structure of the interfacial watermolecules take a multiple changes, and finally takes a hydrophobic
contact layer is formed − 0.81 V vs. SHE.

raising increased demand for a comprehensive and practicalinterface theory. Our recent activity has been directed
towards establishing such theory.
At the interface H+ or OH − ion is adsorbed accompanying an electron-transfer. This so-called redox reaction
is influenced by macroscopic factors including the bias
potential, pH, counter ion density, and the electric double
layer as well as macroscopic factors such as local bondmaking/breaking with the electrode, tunneling/hopping ofelectrons and protons, and hydrogen-bond network. Thisis the essence of electrochemical reactions and relating
the microscopic and macroscopic parts comprehensively
is the target of the research. Lack of information regarding
the interfacial structure and dynamical behavior has been
the primary obstacle to our understanding. In this context,
we have started by simulating the interface of a bare metal
surface and pure water under the finite bias condition.
To approach the problem we have previously developed
the effective screening medium (ESM) based first-principles
molecular dynamics (FPMD) method, which is able to apply
a finite bias to interfaces consisting of the metal slab modeland water molecules [1]. The scheme is applied to theinterface under a negative bias condition ( − 0.7 V vs. thestandard hydrogen electrode (SHE)) to find that an additionally introduced hydronium ion (H3O+) is attracted to thePt electrode and is adsorbed accompanying a discharging
process within the simulation time (7 ps) [2]. See Figure 1
for the snapshots. This is the first simulation of the mostfundamental electrochemical reaction, which is known
as the Volmer step (H3O+ + e − → H(ad) + H2O) of thehydrogen evolution reaction, and suggests important rolesof the structure and response of the water molecules.
The structure of the interfacial water is further investigated to clarify a change with the bias potential. Theinterfacial water molecules take a multiple step to change
the structure from the "H-up" to the "H-down" as the watermonolayer does on the surface, until a hydrophobic contact
layer is formed [3]. The hydrophobic nature can be seen in
the dramatically reduced oxygen atom density in the layer
distribution function. (See the distribution at around 5 Å in
Figure 2)
The simulation was too simplified to be realistic
although suggestive. In the second step we are now trying
to improve the ESM-FPMD to incorporate the pH andcounter ion effect as well as the tunneling of the electrons.
Algorithms to simplify the simulation is also undertaken to

ISSP Activity Report 200711

overcome the limitations in the computational cell size andthe simulation time. Once these improvements are done, wewill then be able to greatly advance our understanding of
the electrochemistry which has attracted chemists over 230
years since the first discovery, expanding the field of thesurface/interface science.
References[1] M. Otani and O. Sugino, Phys. Rev. B 73, (11) 115407 (2006).
[2] M. Otani, I. Hamada, O. Sugino, Y. Okamoto, and T. Ikeshoji,
J. Phys. Soc. Jpn. 77, 0248021 (2008).
[3] M. Otani, I. Hamada, O. Sugino, Y. Okamoto, and T. Ikeshoji,Phys. Chem. Chem. Phys. to be published.
Authors
O. Sugino and M. Otani

Phase Diagram of Dissipative DoubleWell Systems
Kato Group
Quantum tunneling between two states subject to dissipation has been one of central problems in both physics
and chemistry for several decades. A number of studies
on dissipative quantum systems have revealed that dissipation strongly affects the quantum nature of the system.
For example, destruction of quantum superpositions has
been studied extensively in a dissipative two-state model(a spin-Boson model) [1], which is obtained by truncationfrom a double-well system. Further, in the presence ofstrong dissipation, quantum tunneling between two states
is completely suppressed, and a particle is localized in one
well.
Truncation to a two-state model is justified in the large
barrier limit (g = ħω/V<<1), where V is a barrier height andω is a frequency of small oscillation around the minima.
If the barrier height V is comparable to ħω, the doublewell system is never mapped to a two-state model. Do thecharacteristic features of the dissipative two-state systems(e.g. decoherence, localization, etc.) survive in the presenceof large quantum fluctuation? We have studied this problem
by using the path-integral Monte Carlo (PIMC) method [2, 3].
The conventional PIMC method applied to dissipative
double-well systems has several difficulties. Most difficult problem is that the simulation time is proportional to
L2logL, where L is a Trotter number (the division numberof the path along the imaginary time). This makes it hard totreat the low-temperature region corresponding to a large
L. We have developed a new algorithm of the PIMC for
efficient local update. Our idea is based on a decomposition of the weight of the nonlinear potential term to several
Gaussian distributions. In our algorithm, the simulation timeis proportional to LlogL. Our method also reduces the correlation time along the Monte Carlo step. For example, toobtain 1000 independent samples for L=214 on a standard
desktop personal computer, it takes about one month in theconventional method, while only 5 minutes in our improved
method. To overcome critical slowing down, we also introduced cluster algorithm proposed by Werner and Troyer [4].
The localization transition occurs at α=1 in the two-state
limit (g → 0), where α is a dimensionless dissipation strength.
This belongs to the Kosterlitz-Thouless transition [1]. Also
for a finite g, it is expected that the transition belongs to thesame universality class because of symmetry of the model.
Therefore, we can analyze the transition by a finite-size
scaling on the helicity modulus Ψ =(α/2N)Σi,j<xixj>, wherexi is a discretised path. The upper curve of Fig. 1 shows
the localization transition line determined by the PIMC
simulation up to L=214. The coherent-incoherent crossover is defined in several ways. Here, we determine it bydisappearance of a peak in the structure factor S(ω), which
is obtained from the correlation function with the Padé
approximation. The lower curve of Fig. 1 shows the crossover line.
As seen in the phase diagram, both the crossoverand transition line shift towards a large α for a finite g.
Furthermore, the tendency of saturation is clearly observed
for larger values of g. These behaviors indicate that even for
the infinitesimal potential barrier (V << ħω), a finite strengthof dissipation is sufficient for decoherence and localization. Our results also indicate that the basic phenomena
discussed in the dissipative two-state system can beobserved also in the general dissipative double-well systems
even for a small potential barrier.
Our improved algorithm can be applied to otherproblems with continuous variables. One typical example isthe ϕ4 chain. Our idea of Gaussian decomposition suggeststhat such a model can be mapped approximately intoanother model only with discrete variables after integratingout the continuous variables. This may give a new insight
into the original model.
References[1] A. J. Leggett, S. Chakravarty, A. T. Dorsey, M. P. A. Fisher, A.
Garg, and W. Zwerger, Rev. Mod. Phys. 59, 1 (1987).
[2] T. Matsuo, Y. Natsume, and T. Kato, J. Phys. Soc. Jpn. 75,
103002 (2006).
[3] T. Matsuo, Y. Natsume, and T. Kato, Phys. Rev. B 77, 184304(2008).
[4] P. Werner and M. Troyer, Phys. Rev. Lett. 95, 060201 (2005).

Fig. 1. Phase diagram of dissipative double-well systems. The dimensionless parameter g = ħω/V expresses strength of quantum fluctuation, and α is a dissipation strength. The full square shows the phasetransition determined by our simulation with an accuracy of lessthan 1%. The full circle shows the coherent-incoherent crossover
determined from the dynamical structure factor, where their statistical errors are typically given by the symbol size. The empty squareand circle show the known transition and the crossover point in thelarge-barrier limit, respectively.

12

ISSP Activity Report 2007

Authors
T. Matsuoa, Y. Natsumea, and T. Kato
aChiba University

Stripe Phase of Two-Dimensional HoleGas at Half-Filled Higher Landau Levels
Iye Group
When a two-dimensional electron system (2DES) is
subjected to a strong perpendicular magnetic field, electrons
execute cyclotron orbital motion. Quantum mechanically,
the cyclotron radii are quantized to discrete values andthe corresponding energy spectrum is a series of discrete
Landau levels (LLs). In such a situation the effect ofCoulomb interaction becomes particularly crucial. This isreflected to the rich variety of the ground states of 2DES
under a quantizing magnetic field depending on the LL
filling factor ν = nh/eB. The strongly correlation gives riseto fractional quantum Hall liquid states and Wigner crystal
states at lower LL fillings. At half-filling of higher LLs, suchas ν = 11/2, 9/2, and 7/2, the 2DES was found to develop
outstanding anisotropy in resisitivity in spite of no obvious
source of anisotropy [1]. This was subsequently called
stripe phase, which is a sort of charge density wave state,
namely spontaneous development of spatial regions withdifferent LL fillings. The stripe phase is sensitively affectedby an in-plane magnetic field [1] or by an artificial potential
modulation [2].
The previous studies were mostly focused on the 2D
electron systems. Comparative studies on two-dimensionalhole system (2DHS) having different parameter values from
2DES (e.g. higher effective mass) should shed light on thephysical essence of the phenomenon. Although there werea few studies addressing to the stripe phase in 2DHS, thesamples used in those studies were grown on the (113)A
surface of GaAs substrate with Si dopant, so that they had
inherent crystallographic anisotropy in the 2DHS plane.
In the present work, we used 2DHS samples grown on
the (001) substrate with Be dopant [3]. Figure 1(a) showsmagneto resistance of a 2DHS sample for two different
current directions. The QH transition peaks exhibit distinct
anisotropy at half-filling of higher LLs, such as ν = 11/2, 9/2,
and 7/2 (but not at ν = 5/2), indicating the occurrence ofstripe phase at these fillings. Figure 1(b) shows the temperature dependence of the peak height at ν = 7/2 revealing thatthe stripe phase disappears at around 150 mK. The effectof an in-plane magnetic field on the anisotropy is consistent
with the behavior of the stripe phase elucidated for 2DES.
However, the effect is found to be weaker than the case of
2DES, suggesting a stronger orientational pinning mechanism for the hole stripe to the underlying crystalline structure.
References[1] M. P. Lilly et al., Phys. Rev. Lett. 82, (1999) 394; R. R. Du et al.,Solid State Commun. 109, 389 (1999).
[2] A. Endo and Y. Iye, Phys. Rev. B 66, 075333 (2002).
[3] H. Takado, Y. Hashimoto, A. Endo, S. Katsumoto, and Y. Iye, J.
Phys. Soc. Jpn. 76, 074712 (2007).
Authors
H. Takado, Y. Hashimoto, A. Endo, S. Katsumoto, and Y. Iye

Quantum Decoherence inSuperconducting Charge Qubit
Katsumoto Group
Quantum decoherence is one of the biggest mysteriesin quantum mechanics. Several mechanisms have been
proposed for a system to decohere. Most interesting is theone that predicts a "phase transition" between quantum
mechanical region and classical region. Superconductorinsulator transitions in Josephson array systems wereclaimed to be indirect evidences of the mechanism.

Fig. 1 (a) Magnetoresistance traces of a 2DHS sample depicted in theinset. The two traces represent resistivity along two different current
directions. (b) Temperature dependence of the peak resistivity atν = 7/2 for the two directions. The green data represent their ratio.

Fig. 1. Color plot of the source-drain current Isd through the Cooper
pair box against the plane of the gate voltage Vg and the source-drain
voltage Vsd.

ISSP Activity Report 200713

relation is obtained from the microwave frequency dependence of the split width.
Figure 2 shows the microwave frequency as a functionof split width for the device with and without two-dimensional electron system. Apparently the anti-crossing width
is diminished for the device with two-dimensional electrons
and the result evidences quantum decoherence due to theinteraction between the Cooper pairs and the two-dimensional electrons.
Authors
S. Katsumoto, E. Abe, and Y. Kimura

Superstructure-dependent Decoherenceof Surface Electrons on Ge (001)Fig. 2. Applied microwave frequency as a function of split PAJQP
peak position measured from the original JQP peak. Blue and red
circles correspond to the cases with and without two-dimensional
electrons.

However such transitions are consequences of so complicated many-body interactions and far from simple, soundexamples. Successful observation of coherence oscillation of
charge qubit in a Josephson junction system was benefited
from lacking of kinetic freedom in the superconducting
energy gap. This situation, which is an ideal stage forquantum information processing, at the same time can bea test stage of quantum decoherence mechanism. Here we
report tuning of quantum decoherence through the environment kinetic freedom.
The energy gap region is like a blank stage with qubit
as the single player. The qubit in the present experiment iscomposed of a "single Cooper pair box", two Cooper pairreservoirs (source and drain) and two Josephson junctions,which connect these each reservoir and the box. In the
single Cooper pair box, the number of Cooper pairs isquantized with the Coulomb energy. One of the Josephson
junctions is composed as an interferometer so that theJosephson energy can be tuned by external magnetic field.
All of these structures are made of vacuum evaporated
aluminum. Now we introduce the other players. The qubit
device is placed on an AlGaAs/GaAs heterostructure, towhich two dimensional carriers can be introduced by lightillumination. Such free carriers interact with Cooper pairs
in qubit through Coulomb interaction. With increasing
the number of electrons, the number of kinetic freedom
increases and at the critical point the transition to classicalregion is expected.
The quantum coherence can be most easily and clearlyobserved in the frequency domain, where energy levelanti-crossing due to the coherence between two different
charge states is observed. For such experiment, we utilized
current peaks due to Josephson-quasiparticle (JQP)
tunneling, in which a Cooper pair tunnels through one of
the junctions and two quasi-particles tunnel through theother. Figure 1 shows color plot of the current through thedevice as a function of the gate voltage and the sourcedrain voltage. White colored high current regions are the
JQP peaks. With applying microwave, these peaks split into
two or three due to the photon-assisted (PA) JQP process.
The PAJQP process is caused by microwave excitationfrom lower energy level to upper thus a kind of dispersion

14

ISSP Activity Report 2007

Komori GroupElectron loses its phase memory when it is inelastically scattered. The way of decoherence depends on
the scattering mechanism. On metal and semiconductor
surfaces, the observed decay of electronic standing waves
is attributed to the decoherence, and is determined byelectron-phonon or electron-electron scattering. We havestudied the decoherence of the surface state electrons at
the clean Ge(001) surface, where two stable superstructureswith almost the same quasi one-dimensional (1D) electronic
state coexist at 80 K.
The surface consists of buckled Ge dimers, which alignin the [110] direction and form dimer rows. The empty
π*-band made of the dangling bond states of the surface Ge
dimers behaves like a 1D free electron along the dimer row.
There are two stable superstructures, p(2×2) and c(4×2) (seeFig. 1), and we can control the surface structure reversibly between them with hysteresis by changing the sample
bias voltage VB for the surface observation using scanningtunneling microscopy (STM). [1] We measured the decayof the π*-band standing waves from monoatomic steps in

Fig. 1. Topographic (a) and dI/dV (b) images of a clean Ge(001)surface at 80 K. Two superstructures, c(4×2) and p(2×2) shown
schematically in the upper panels, were intentionally positioned at
the lower terrace of the step by a surface manipulation [1]. Standing
waves of the π* electrons from the step edge are seen in (b).

Fig. 2. Cross sections of the standing waves in the dI/dV images forc(2×2) (a,c) and p(4×2) (b,d) areas. [1] The sample bias voltage VB is
0.5 V (a,b) or 0.4 V (c,d). The wave length decreases with increasing
VB. Blue and red curves are damping oscillation curves fitted tothe data. The decay length of the wave for the same bias voltage isshorter on the c(4×2) surface than on the p(2×2) surface.

differential conductance dI/dV images after preparing asurface with the both superstructures by the VB change.
Figure 1(a,b) shows the topographic and dI/dV imagesof the same area. On this surface, both p(2×2) and c(4×2)
structures coexist on the lower terrace of a single step [1].
We can see that the sanding waves from the step decay inseveral nm in Fig. 1 (b). The cross sections of the differential conductance image from the step are plotted in Fig.
2 for VB = 0.4 and 0.5 V on the p(2×2) and c(4×2) areas.
We fitted each standing wave to the equation: dI/dV(x) =Aexp(-2x/Lc)sin(2kx + δ). Here the fitting parameter Lc isthe decay length of the wave, A and δ the amplitude andphase shift of the wave, and k the wave number of the π*electron. The dispersion relation of the π*-band for eachsurface obtained by this analysis is almost common to theboth surfaces. Whereas, Lc on the p(2×2) surface is longer
than on the c(4×2) for both bias voltages. This suggests thatthe superstructure-dependent electron-phonon interactionin the two surfaces makes the decoherence of the surface
electrons different.

Fig. 1. IRAS spectra of NO on Pt(997) as a function of substratetemperature. The coverage of NO is about 0.08 ML.

migrates on the surface from the first impact point to apossible adsorption site. At 11 K, the molecules are trappedat four adsorption sites on Pt(997): the on-top sites on the(111) terrace (OT), the hollow sites on the (111) terrace(HT), the bridge sites at the step (BS) and the hollow sites atthe step downstream (HS), which have been evidenced byinfrared reflection absorption spectroscopy (IRAS) (Figure 1)[1] and first-principles calculations [2].
(a)

(b)

Reference[1] Y. Takagi, Y. Yoshimoto, K. Nakatsuji, and F. Komori, J. Phys.
Soc. Jpn. 72, 2425 (2003); Surf. Sci. 559, 1(2004); Phys. Rev. B 75
115304 (2007).
Authors
K. Nakatsuji and F. Komori

Microscopic diffusion processes
of NO on the Pt(997) surface
Yoshinobu GroupThe microscopic diffusion processes of NO molecules
on Pt(997) at low coverage were investigated using
time-resolved infrared reflection absorption spectroscopy(TR-ITAS). When NO molecules absorb on the Pt(997)
surface at low temperature, each molecule transiently

Fig. 2. (a) A series of TR-IRAS spectra of NO on Pt(997) after an
NO gas injection. (Coverage: 0.02 ML, Temperature: 76 K), (b) The
change of the fractional coverage of each NO species in Figure2(a) as a function of the elapsed time. Dots: OT species, cross: HT
species, triangles: BS species, lines: fitted lines based on the firstorder kinetics model.

ISSP Activity Report 200715

Fig. 3 Probable diffusion routes to the bridge site of the upper step.
Route A: from the upper terrace, Route B: from the lower terrace.
The starting point is set at an on-top site in this figure as an example.
*: These adsorption energies are taken from Ref. [2].

By heating the surface to 45 K, the HS species migrate
up to the BS sites; the migration barrier is roughly estimatedto be 120 meV. In the temperature range from 70 K to 77 K,
TR-IRAS measurements were carried out to observe the sitechange of OT species to the adjacent HT sites at isothermal
conditions (Figure 2). Based on the Arrhenius analysis oftime-resolved IRAS data at several temperatures, the activation barrier and the pre-exponential factor are estimated to
be 200 meV and 2×1011 s-1, respectively.
In the temperature range from 100 K to 110 K, the HT
species migrate across the terrace and finally reach the BS
sites. The activation barrier between the HT sites and its
pre-exponential factor are estimated to be 290 meV and6.5×1011 s-1, respectively, form the TR-IRAS data together
with kinetic Monte Carlo simulations.
On the whole, the quantitative microscopic picture of
NO migration on Pt(997) has been established. Figure 3
summarizes the schematic diagram of NO on Pt(997) andthe potential energy surfaces across the possible migrationroutes.

effects in situ using a low-temperature scanning tunnelingmicroscopy / spectroscopy.
Figure 1(a) show a 9 monolayer-thickness (ML) Pb
island whose size is a few times larger than the coherent
length (ξ). Tunneling spectra showing a superconducting
gap were taken on this island under various magnetic fields
perpendicular to the surface plane. It turned out that thegap, becoming shallow with the magnetic field [1,2], shows
spatial dependence in the island.
In order to elucidate the observed site dependence, we
made an image of the tunneling conductance at the bottomof the gap (zero bias conductance; ZBC) The ZBC image
taken at zero field (b) indicates homogenous superconductivity. At 0.5 T (c), however, ZBC around the peripheral
region is a little bit larger than the center due to the field
penetration. Then, at 0.6 T (d), only the center region has
high ZBC due to a vortex formation, which was confirmedby an agreement of its profile with the result of linearlized
Ginzburg-Landau (LGL) equations (f). Finally, at 0.9 T (e) allarea has high ZBC, indicating complete breaking of superconductivity in all the area.
Visualization of the vortex formation and the breakdownof superconductivity by STM/STS enables us to determine
critical fields of the vortex penetration/expulsion (Hp1/Hex1)
and the complete breakdown of superconductivity (HSN).
We measured the critical fields on various sizes of islands.
The results taken on 9ML islands are plotted in Figure 2. It
turned out that there is a critical island size for the vortex
formation. Larger islands tend to have lower vortex penetration/expulsion and breakdown fields. According to theoretical studies based on LGL equations, the islands whose
radius is smaller than 1.3 times coherent length do not form

References[1] N. Tsukahara, K. Mukai, Y. Yamashita, and J. Yoshinobu, Surf.
Sci. 600, 3560 (2007).
[2] N. Tsukahara, K. Mukai, Y. Yamashita, J. Yoshinobu, andH. Aizawa, Sur. Sci. 600, 3477 (2007).
[3] N. Tsukahara, K. Mukai, Y. Yamashita, and J. Yoshinobu,
J. Chem. Phys. 128, 054701 (2008).
Authors
N. Tsukahara, K. Mukai, Y. Yamashita, J. Yoshinobu, and H. AizawaaaHokkaido University

Superconductivity of Nano-size Island
Structures Studied by Scanning
Tunneling Microscopy/ Spectroscopy
Hasegawa Group
Superconductors whose size are comparable to as small
as their coherence length exhibit properties quite different
from those of bulk counterparts. In order to study them inreal space, we first prepared nanosize flat-top single-crystalline Pb island structures on a Si substrate, whose typicalsizes are 25-75 nm in radius and 3 nm in thickness, andthen investigated their superconducting properties and size

16

ISSP Activity Report 2007

Fig. 1. (a) STM image of a 9ML Pb island showing an area where thefollowing ZBC images were taken. (b)-(e) ZBC images taken under themagnetic field of 0.0 T (b), 0.5 T (c), 0.6 T (d) and 0.9 T (e). (f) crosssectional ZBC profiles. A line drawn for the 0.6 T data is a spatial
profile of the order parameter calculated from the LGL equations.

Fig. 2. Lateral size dependence of various critical fields; the transition field between the superconducting and normal states HSN, vortex
penetration Hp1 and vortex expulsion Hex1. All critical fields wereobtained from 9ML-thickness Pb islands.

vortices. From the critical size, the coherence length of the
9ML islands was estimated to be 31+3 nm. As the radiusincreases HSN approaches to that of thin films; namely, thecritical field for the surface superconductivity Hc3, which
is given by 1.695 × h/4πeξ2. The value of Hc3 calculated
from the estimated coherence length is marked in the plot.
Although this analysis neglects an effect of island shape, theexperimental results are quite consistent qualitatively withthe results of the theoretical studies.
References[1] T. Nishio, M. Ono, T. Eguchi, H. Sakata, and Y. Hasegawa, Appl.
Phys. Lett. 88, 113115 (2006).
[2] T. Nishio, M. Ono, T. Eguchi, H. Sakata1, and Y. Hasegawa, Jpn.
J. Appl. Phys. 46, L880 (2007).
Authors
T. Nishio, T. An, A. Nomura, K. Miyachi, T. Eguchi, H. Sakataa, and
Y. Hasegawa
aDept. of Physics, Science Univ. of Tokyo.

Fig. 1. The band structure of SrTiO3 close to the conduction band(CB) bottom. The intrinsic Fermi level is located about 100 eV below
the CB edge. In a field-effect transistor (Inset) the Fermi level of
SrTiO3 can be modulated by applying a positive or negative bias to
the gate electrode.

(Fig. 1), the Fermi level can be pushed higher or lower by
accumulating or depleting carriers from the device channel.
The distance between the Fermi level and the CB edge
can be measured by determining the activation energy forthermally activated conduction in the transistor channel.
The temperature dependence of channel current in apartially oxygen-depleted SrTiO3 device is shown in Fig. 2.
Particularly in the depletion regime, where the gate bias
is negative, the channel current shows a strong temperature dependence. By measuring the channel current as afunction of temperature at a fixed gate bias, the activationenergy can be found from an Arrhenius plot. Since thenumber of carriers depleted or accumulated in the FETchannel (and thus the position of the Fermi level) is a linear
function of the gate bias and gate insulator capacitance, it
is possible to estimate the density of states as a function ofenergy (from the CB bottom).
The plot in Fig. 3 shows a linear increase in EA when thegate bias is close to zero, corresponding to approximately

Mapping Defect States in SrTiO3
Lippmaa Group
Despite the seemingly simple crystal structure, theelectronic properties of SrTiO3 continue offer surprises,
such as metallic and even superconducting interfaces in
SrTiO3 heterostructures. SrTiO3 is an unusual material in
the sense that transitions from the band insulating state tometallic and even superconducting states occur at very low
carrier densities, at around 1018 and 1020 cm-3, respectively.
While this feature suggests several possible avenues for
doping experiments, such as charge transfer or field effect,
it also means that SrTiO3 is prone to inadvertent doping
by a very low density of defects, such as oxygen vacancies. Since the Fermi level in intrinsic SrTiO3 is pinned very
close to the conduction band edge, it is more accurate toview SrTiO3 as a wide-gap semiconductor (Eg = 3.2 eV),rather than a classical band insulator. In intrinsic SrTiO3,
the Fermi level is about 100 meV below the conduction
band (CB) bottom. In a SrTiO3-based field-effect transistor

Fig. 2. Channel current of a SrTiO3 FET as a function of gate bias
and temperature. In depletion mode (negative gate bias) the channel
current is a strong function of temperature due to thermally activatedtransport in localized in-gap states close to the conduction bandedge.

ISSP Activity Report 200717

Fig. 3. Activation energies determined from the temperature dependence of channel current for each gate bias. At Vg = +0.5 V, thedevice becomes metallic, i.e. the Fermi level is pushed across themobility edge. For negative bias, the activation energy increases
linearly (constant density of states) until saturating at 0.3 eV, whichcorresponds to the start of an in-gap impurity band related to oxygen
vacancies in SrTiO3.

Fig. 1. O(N) calculation time of the conductance of the (5,5) carbon
nanotubes with various system lengths on a Hitachi SR11000 in
ISSP. The solid and broken lines represent the calculation time of the
Chebyshev polynomial development and the split operator methods,
respectively. In the inset, we show the calculation time for the shortsystem lengths [2].

constant density of states. An abrupt saturating behavior isseen below about −0.5 V, where the activation energy does
not increase beyond 0.3 eV. This shows that in oxygendeficient SrTiO3, there is a large density of states at about
0.3 eV below the CB bottom. We assign this state to ogygenvacancies in SrTiO3.
Reference[1] K. Shibuya, T. Ohnishi, T. Uozumi, T. Sato, K. Nishio, and M.
Lippmaa, Appl. Phys. Lett. 92, 032109 (2008).
Authors
M. Lippmaa, K. Nishio, T. Ohnishi, K. Shibuyaa, T. Uozumib, T. SatocaJülichbCanoncToshiba

Quantum Transport Simulations of Carbon
Nanotube Field-Effect Transistors
Using O(N) Time-Dependent Approach
Kobayashi Group
Toward developments of nanoscale devices, intensive
works have been performed in the investigation of the transport properties of nanoscale systems. Carbon nanotubes
show various electronic states and are expected to be a
candidate element of nanoscale devices. We have investigated the quantum transport properties of carbon nanotube
field-effect transistors using time-dependent wave-packet
diffusion method [1,2]. We have succeeded in O(N)calculations for transport of large systems with micronorder lengths using an atomistic model and investigationfrom ballistic regime to diffusive regimes based on the
Kubo-Greenwood formula.
We have investigated the effects of electrode contact
and electron-phonon coupling on the quantum transport properties of carbon nanotube devices with micronorder channel lengths. We have elucidated the electron
mobility of a (10,0)-semiconducting carbon nanotube fieldeffect transistor with several μm channel lengths for the

18

ISSP Activity Report 2007

Fig. 2. Calculated time-dependent diffusion coefficient of the (10,0)
carbon nanotube field-effect transistor with 1 μm channel length forvarious gate voltages Vg [1].

applied gate voltages in the absence of phonons, vacancies,
and impurities. Furthermore, we investigated the phonon
scattering effects on the mobility and conductance of the
carbon nanotubes. The longitudinal acoustic phonon andoptical phonon modes with a small momentum were considered at room temperature. The conductance of a (5,5)
metallic carbon nanotube is decreased around the Fermi
energy by the optical phonon scattering. On the other hand,
the mobility and conductance of the (10,0)-semiconducting
carbon nanotube are decreased by the acoustic phononmode. We have also elucidated the transitions of transport
behaviors of nanotubes from ballistic to diffusive regimes.
References[1] H. Ishii, N. Kobayashi, and K. Hirose, Phys. Rev. B 76, 205432(2007)[2] H. Ishii, N. Kobayashi, and K. Hirose, Appl. Surf. Sci. (2008) inpress.
Authors
H. Ishiia, N. Kobayashib and K. Hirosec
aUniversity of Tsukuba
bISSP and University of Tsukuba
cNEC Corporation

Two Dimensional Anti-ferromagnetic
Solid ３He in High Magnetic Fields
Ishimoto Group
３He

film adsorbed on graphite surface is one of theideal two-dimensional (2D) Fermi systems that can realize
not only a Fermi liquid but also a quantum solid just bychanging the adsorption density. The transition between
liquid and solid is recently considered to be a Mott transition [1]. The 2D solid form a quantum spin magnet with
nuclear spin S = 1/2 on a triangular lattice. Because of
the hard-core potential between ３ He atoms, the higher
order multiple spin exchange (MSE) processes as well as
two-particle exchange play an important role. The exchangeof even number of particles is anti-ferromagnetic (AFM),
while that of odd ones is ferromagnetic (FM). The competition among them, in addition to the geometrical frustration
inherent to a triangular lattice structure, makes AFH solid
３He film strongly frustrated.
According to recent theoretical studies in the exact
Fig. 2. Magnetization curve for the AFM 4/7 phase at 0.7 mK. The
vertical axis is normalized by the saturation magnetization.

diagonalization method for this finite size system, noneof long range ordered AFM phases survives even at T =0 K due to quantum fluctuations. The exchange interaction parameters obtained at high temperatures for the AFM
phase ( so called a 4/7 phase ) suggests a quantum spin
liquid (QSL) ground state with a finite excitation gap [2], anda magnetization plateau. Experimentally the heat capacity
and susceptibility measurements indicate a gapless spin
liquid ground state [3][4]. Our previous measurement witha double–gradient Faraday-type magnetometer suggests theexistence of magnetization plateau in the AFM 4/7 phaseon 4He mono layer pre-plated graphite. To resolve such a
discrepancy, we investigated the magnetization curve byusing NMR technique over the wide frequency range from
7 MHz to 360 MHz in a single experimental run.
Figure 1 shows the magnetization of AFM 4/7 phase on
4He mono layer pre-plated graphite as a function of temperature for various magnetic fields. Here the vertical axis is
the polarization estimated from the saturation magnetization obtained by fitting the temperature dependence in thehigh temperature region. Figure 2 shows the magnetization
curve obtained from Fig.1. It shows a clear plateau at halfof saturation magnetization, followed by its full saturation
at about 10 T. It is essential to take account of five and six
multiple spin exchange interactions to explain the results.
References[1] A. Casey, H. Patel, J. Nyʼki, B. P. Cowan, and J. Saunders, Phys.
Rev. Lett. 81,115301(2003).
[2] G. Misguich, B. Bernu, C. Lhuillier, and C. Waldtmann, Phys.
Rev. Lett. 81,1098 (1998).
[3] K. Ishida, M. Morishita, K. Yawata, and H. Fukuyama, Phys. Rev.
Lett. 79, 3451 (1997).
[4] R. Masutomi, Y. Karaki, and H. Ishimoto, Phys. Rev. Lett. 92,
025301 (2004).
Authors
H. Nema, A. Yamaguchi, T. Hayakawa, and H. Ishimoto

Fig. 1. Temperature dependence of magnetization in the AFM 4/7
phase for various fields. The vertical axis is normalized by the saturation magnetization. The lines in the inset are the Brillouin functions.

ISSP Activity Report 200719

First Evidences of Vortex Fluid toSupersolid Transition in Solid 4He
Kubota GroupThe question, whether particles forming a solid lattice,
can show also properties of superfluid at the same time, hasa long history. From the first theoretical study's negative
answer [1] to positive discussions [2], the original idea of
possible superfluidity was discussed based on Bose-Einsteincondensation(BEC) of particles themselves or imperfectionsin the system of a quantum solid. The latter refers to either
vacancies or interstitials and dislocations or all together, but
in practice only the simplest vacancies and interstitials had
been evaluated and from estimated densities of imperfections in solid 4He the occurrence of such a state, now called"Supersolid" had been expected to be at very low temperatures. The situation has been changed drastically since
the discovery of so-called "non-classical rotational inertia"(NCRI) in solid 4He, by Kim and Chan [3]. NCRI is common
in all known other superfluids and it had been suggested tobe observable by Leggett[2-c] long time ago for a supersolid.
It then became a problem since reported "onset temperature, To" was higher than 150 mK or 300 mK, and not muchdifferent (within a factor of 10) from liquid 4He λ line (2.171.743K, depending on pressure), where superfluid transition
is primarily governed by BEC of 4He atoms. P. W. Anderson[4], then proposed that the reported "NCRI" in solid He
is actually "non-linear rotational susceptibility (NLRS)" of(a)

vortex fluid state and he pointed out that real Τ c of supersolid transition should exist at some lower temperature.
Due to the very stable torsional oscillator conditions[5] and the world leading rotation cryostat [6], which wasconstructed at the ISSP, we are able to report and confirm
the following points:1). From detailed AC velocity dependent responses of torsional oscillator (TO) containing solid
4He samples we could determine uniquely an onset temperature To, which is much higher than so far reported data.
We discuss the appearance of vortex fluid state below To [7].
2). An indication of transition from vortex fluid state to thereal 3D supersolid state is discovered below Tc = ~75 mK[8]. 3). Existence of vortex lines though the solid 4He sample
under DC rotation below Tc is confirmed by the linear
dependence on DC rotational speed Ω, Fig.1(b) [9]. We cancomment on possible misinterpretation in the recent reportby Day and Beamish on the "hardening" of the Solid 4He[10]. For recent reviews we find [11, 12].
References[1] Penrose and Onsagar, Phys. Rev. 104, 576 (1956).
[2] a). Reatto and Chester, Phys. Rev. 155, 88 (1967); Chester, Phys.
Rev. A 2, 256 (1970); b). Andreev and Lifshitz, JETP 29, 1107
(1969); c). Leggett, Phys. Rev. Lett. 25, 1543 (1970).
[3] Kim and Chan, Nature 427, 225 (2004); Science 305, 1941
(2004); Phys. Rev. Lett. 97, 115302 (2006).
[4] P. W. Anderson, Nature Phys. 3, 160 (2007).
[5] M. Fukuda et al., J. Low Temp. Phys. 113, 423 (1998).
[6] M. Kubota et al., Physica B 329-333, 1577 (2003).
[7] Penzev, Yasuta, and Kubota, arXiv:0711.0212, submitted to PRL.
[8] Penzev, Yasuta, Shimizu, Gumann, and Kubota, submitted to
LT25.
[9] Shimizu, Penzev, Yasuta, Gumann, and Kubota, submitted toLT25.
[10] Day and Beamish, Nature 450, 853 (2007).
[11] N. Prokof'ev, Advances in physics, 56, 381 (2007).
[12]Balibar and Caupin, J. Phys. Cond. Mat. 20, 173201 (2008).
Authors
A. Penzyev, N. Shimizu, Y. Yasuta, P. Gumann, and M. Kubota.

Interlayer Magnetotransportin Layered Conductors:
Multilayer Dirac Electron Systems
and Magnetic Breakdown Systems
Osada Group(b)

Fig. 1. (a) TO Energy dissipation change ΔQ-1 vs DC angular velocity
Ω. Extra dissipation appears over a temperature range T < Tc=~75-85 mK below a dissipation peak (round peak) and the linear
dependence to Ω at constant temperatures is more clearly seen in(b) and it gives clear evidence of vortex lines penetration, in contrast
to a possible Ω2 dependence expected to be caused by solid vibrations or other solid properties under DC rotation.

20

ISSP Activity Report 2007

Interlayer magnetotransport in two multilayer electron
systems has been investigated experimentally and theoretically.
(1) Dirac electron systems
We have considered the interlayer magnetotransport
in multilayer Dirac electron systems, in which two-dimensional electron layers with Dirac cone dispersion stack withweak interlayer coupling. The Fermi level is assumed to
be at the contact point of Dirac cones, so that the systems
become zero-gap conductors. Under magnetic fields, then = 0 Landau level always appears at the contact point (zeromode), and its overlap with other Landau levels vanishes atrather weak field. In this quantum limit, the interlayer transport is dominated by the tunneling between zero-modes
on neighboring two layers. The increase of zero-mode
degeneracy causes strong negative interlayer magnetoresistance under normal magnetic fields. Spin splitting of thezero-mode could change the negative magnetoresistanceto positive in high magnetic fields and low temperatures.

When parallel component of magnetic field is additionallyapplied, the interlayer resistance recovers from negative
magnetoresistance reflecting the decrease of overlap of
two zero-modes between which tunneling is allowed. The
present model has successfully explained the magnetotransport features observed in an organic conductor α-(BEDTTTF)2I3 by Tajima and Kajita. This fact strongly supports therealization of a Dirac electron system in α-(BEDT-TTF)2I3
predicted theoretically.

(2) magnetic breakdown systems
We have studied the change of angle-dependent magnetoresistance oscillation phenomena through magnetic
breakdown process. Using the double-axis rotator, wehave measured the angle-dependent pattern of interlayer
magnetoresistance of an organic conductor α -(BEDTTTF)2KHg(SCN)4 over whole magnetic field orientation
as a function of temperature, and obtained complicatedmixture of various angular effects. The electronic system
of CDW phase of α-(BEDT-TTF)2KHg(SCN)4 can be considered as the quasi-two-dimensional (Q2D) electron systemwith cylindrical Fermi surface under additional CDW
potential. The CDW amplitude is controlled by temperature. Using quantum tunneling picture, we have shownthat the Kajita oscillation of the original cylindrical Fermi
surface is modified by the Lebed-like resonance withdecreasing temperature, since the magnetic breakdown
through the CDW gap is suppressed. The present modelhas well explained the observed features in α-(BEDTTTF)2KHg(SCN)4.
Authors
T. Osada, K. Uchida, and T. Konoike

Effect of Pressure on the One-dimensional
Organic Conductor (TMTTF)2SbF6
Uwatoko Group

Fig. 1. (a) Dirac cone dispersion and (b) Landau level configurationin the 2D massless Dirac electron system. (c) Interlayer conductivity
in the multilayer Dirac electron system in magnetic fields calculatedusing a lattice parameter of α-(BEDT-TTF)2I3.

Fig. 2. Second derivative pattern of interlayer resistance of α-(BEDTTTF)2KHg(SCN)4 scanned over the whole magnetic field orientation.

The (TMTCF)2X series (C = S and Se, X = monovalent anion) comprises traditional quasi-one-dimensional
compounds exhibiting highly anisotropic electron transfers.
Almost two decades ago, a general pressure temperature(P-T) phase diagram was proposed for (TMTCF)2X families.
The compound (TMTTF)2SbF6 has the largest monovalentanion in the (TMTTF)2X series, which is advantageous toinvestigate the wide range of the P-T phase diagram for asingle system. The ground state of this compound has an AForder. The transition temperature is TN = 8 K and a chargeorder (CO) occurs at TCO = 154 K. Combining these facts
with the concepts embodied in the general phase diagram,the possible sP and SDW phases exist in (TMTTF)2SbF6 at
low pressures, and the SC phase may be stabilized withfurther application of pressure. Therefore, the systematicinvestigation of (TMTTF)2SbF6 has been eagerly anticipated in order to understand the entire electron correlation
system in all the (TMTCF)2X salts.
We obtain the general P-T phase diagram by adding(TMTTF)2SbF6, as shown in Figure 1. In the figure, theshift of pressure axis in (TMTTF)2PF6 and (TMTTF)2AsF6
are estimated to occur at 1 GPa and 0.5 GPa, as indicatedby the arrow. The possible sP, AF, and SDW regions arereferred to the previous general (TMTCF)2X phase diagram.
After the CO and AF phases vanish in (TMTTF)2SbF6, newphase appears in resistivity. The transition temperature
TS ~ 20 K of the SbF6 salt is comparable to the temperature of the sP and SDW transitions in the AsF6 and PF6
salts. A further increase in the pressure probably reduces
the effective on-site Coulomb interaction and breaks thenesting of the Fermi surface. Then, the superconductivity
appears at 5.7 GPa, finally. The SC phase is observed over
the wide pressure range of 5.4 < P ≤ 9 GPa. As comparedto the SC phase of other TMTTF compounds, the pressure
range is anomalously wide. The SC region of PF6 and AsF6

ISSP Activity Report 200721

Photo-induced Metallic State Observedby Terahertz Pump-probe Spectroscopy
Suemoto and Hiroi Groups

Fig. 1. P-T phase diagram of (TMTTF)2SbF6 constructed from thedata obtained in this study. It is superposed on previously obtainedphase diagrams of (TMTTF)2PF6 and (TMTTF)2AsF6. The notationsof CO, AF, sP, SDW, and SC refer to charge order, antiferromagnet,
spin-Peierls, spin-density-wave, and superconducting states, respectively. The data obtained in this study are described as solid green
diamonds (Tρ), as solid light blue circles (TS), and solid blue circles(Tc). The SC regions of the PF6 and AsF6 salts are drawn in the SC
phase of SbF6.

are described as a violet shadow in Figure 1. Moreover,
the SC phase of (TMTTF)2SbF6 has two local maxima andthe SC phase is deeply into the TS phase. With regard to
the double maxima in the SC phase, the heavy fermion
materials CeCu2Si2/Ge2 have been known to depict thisfeature. It could be possible that the mechanism underlying
the formation of the double maxima in (TMTTF)2SbF6 issimilar to that in heavy fermion compounds. To clarify thesuperconducting properties, more detailed investigations are
currently underway.
From the previous reports of the (TMTTF)2X family, theAF ground state (commensurate SDW) appears between
the sP and the SDW states. The AF state corresponds to
the ground state of (TMTTF)2Br at the ambient pressure(TN = 15 K). Based on previously obtained phase diagrams,
an unusual ground state change such as AF(1)-sP-AF(2)SDW-SC can be expected in (TMTTF)2SbF6, but the expectation is not consistent with dimensional effect by pressure.
Recently, 13C-NMR and ESR measurements indicated thatthe characteristic features of the AF state at the ambientpressure in (TMTTF)2SbF6 (AF(1)) and (TMTTF)2Br (that is,AF(2)) are slightly different. While SbF6 has considerable
3-D AF interaction, the nature of AF(2) in (TMTTF)2Br ischaracteristic of a low-dimensional magnet. Furthermore,
the AF wave vector Q = (1/2, 0, 0) for (TMTTF)2SbF6 isdifferent from that (Q = (1/2, 1/4, 0)) for (TMTTF)2Br. It
is natural to consider that the differences in the nature of
the AF phase between (TMTTF)2SbF6 and (TMTTF)2Br are
related to the existence of a clear or nonclear CO phasein paramagnetic region. Therefore, we cannot deny thepossibility of the existence of an AF phase like (TMTTF)2Br(AF(2)) after the disappearing the CO phase. And also, sP
phase may exist between two AF phases in (TMTTF)2SbF6.

Recently, we have been studying the photo-induced
phase transitions by using terahertz (THz) pump-probe
spectroscopy. Here, we present the picosecond dynamics
of the insulator-metal transition induced by photoexcitationin vanadium dioxide (VO2) [1]. VO2 exhibits an insulator tometal phase transition at 340 K accompanying a structurechange from rutile to monoclinic phase. THz pulses, whose
spectral range is from 0.4 to 12 meV, can probe directly theultrafast free carrier response. The dynamical behavior after
photoexcitation at 800 nm in VO2 thin film was investigatedby the optical-pump THz-probe spectroscopy.
Figure 1 shows the temperature dependences of thetransmittance of the THz wave and the electrical resistivity in steady state. The decrease of the transmittancecorresponds to a decrease of the electrical resistivity,that is, an increase of dc electric conductivity. The good
agreement between the transmittance and the resistivity
indicates that the THz probe pulses detect properly thefree carrier response of the metallic state. The transmitted
THz waveforms with optical pump excitation at different
time delays are shown in Fig. 2. After the photoexcitation,
the amplitudes of the THz waves decrease with increasing
the delay time ∆t. In order to investigate the THz response
in detail, the time evolutions of the transient transmittance changes at different excitation densities were evalu-

Fig. 1. Temperature dependences of the power transmittance of theTHz pulse (left axis) and the normalized electric resistivity (right axis).

Reference[1] M. Itoi, C. Araki, M. Hedo, Y. Uwatoko, and T. Nakamura, J.
Phys. Soc. Jpn. 77, 023701 (2008).
Authors
M. Itoia, C. Araki, M. Hedob, Y. Uwatoko, and T. Nakamurac
aAoyama Gakuin University
bThe University of Ryukyus
cInstitute for Molecular Science

22

ISSP Activity Report 2007

Fig. 2. The transmitted terahertz waveforms with the photoexcitation at
different delay times between optical pump and THz probe pulses ∆t.

Fig. 3. The temporal evolutions of the transmittance changes atdifferent excitation densities.

ated as shown in Fig. 3. The temporal profiles show adrastic change around 4.0 mJ i.e., the transmission change
increases significantly above this excitation density. The
decay rates increase from 500 ps to 20 ns with increasing
the excitation density. These transient changes of the transmittance recover in the time interval of the laser pulses of
200 μs.
The time evolution of the transmittance at 4.0 mJ/cm2
in Fig. 3 shows a two-step decrease, a fast onset within 800
fs and a gradual growth after 2 ps. This behavior is very
similar to the temporal profile of the fraction of the hightemperature phase measured by means of the time resolved
X-ray and electron diffraction spectroscopy [2]. This good
agreement indicates that the time profile of the transienttransmittance change, which reflects the conductivity
change, definitely corresponds to the fraction of the hightemperature phase accompanying the change of the lattice
structure. Furthermore, this fact indicates that the photoinduced insulator-metal phase transition is of the Peierls
type.
References[1] M. Nakajima, N. Takubo, Z. Hiroi, Y. Ueda, and T. Suemoto,Appl. Phys. Lett. 92, 011907 (2008).
[2] A. Cavalleri, Cs. Tóth, C. W. Siders, J. A. Squier, F. Raksi,
P. Forget, and J. C. Kieffer, Phys. Rev. Lett. 87, 237401 (2001); P.
Baum, D. S. Yang, and A. H. Zewail, Science 318, 788 (2007).
Authors
M. Nakajima, N. Takubo, Z. Hiroi, and T. SuemotoFirefly Bioluminescence: Quantum
yield and pH-sensitive Color Change
Akiyama Group
Firefly bioluminescence [1] is known as one of themost efficient light-emission systems on the earth having
wonderful color change mechanism with pH, and has
attracted extensive interest by many scientists not onlyin the field of biology but also in chemistry, physics,
electronics, and photonics. The frequently and widely
cited properties of firefly bioluminescence with extremely
high efficiency and pH-dependent emission-color change
between yellow-green and red modeled as chemical equilibrium between two corresponding states have been basedon a seminal report in 1960 [1], which measured a highFig. 1. Quantitative luminescence spectra of firefly bioluminescence(Photinus pyralis) for 2.98 × 1011 luciferin molecules at various pH
values with Tris or GTA buffers. The vertical axis is scaled by anabsolute number of emitted photons in a unit of eV-1. Integrated
area of each spectrum gives total number of emitted photons, or
integrated quanta of luminescence.

quantum yield of 88±25%. However, necessity for morereliable and refined experiments has long been addressed,
because the same group later found and commented that
luciferin has optical isomers and only the D-form contributes to bioluminescence, and that the purity of D-luciferin
used in their measurement was probably significantly
degraded by racemization. In spite of this problem, nocorrection or confirmation had been made. A new quantumyield determination with pure D-luciferin was long overdue.
We quantified quantum yields and emission-color
change of firefly (Photinus pyralis) bioluminescence forpH from 5.4 to 8.9 [2], using our recently developed totalphoton-flux spectrometer [3] for bio/chemiluminescence.
We determined the highest quantum yield of 41.0±7.4%(1 standard deviation estimate, coverage factor k=1) at
pH 8.5. This value is still higher than but closer to thoseof other bio/chemiluminescence systems such as 28% in
cypridina (sea firefly), 17% in aequorin (jelly fish), 30% inbacteria, and 1.24% in luminol. Moreover, we measured
quantitative bioluminescence spectra in the absolute
number of all emitted photons, as shown in Fig. 1, andfound that all the spectra were systematically and quantitatively decomposed into one pH-sensitive and two pH-insensitive Gaussian peaks, and that no intensity conversion
between yellow-green and red emissions but mere intensity
variation of the pH-sensitive Gaussian peak at 2.2 eV causes
the changes in apparent emission colors. These cast serious
doubts on long-standing pictures of bioluminescence-color
determination by the pH equilibrium between yellow-green
and red states.
References[1] H. H. Seliger and W. D. McElroy, Arch. Biochem. Biophys. 88,
136 (1960).
[2] Y. Ando et al., Nature Photonics 2, 44 (2008).
[3] Y. Ando et al., Photochemistry and Photobiology 83, 1205
(2007).
Authors
Y. Ando, K. Niwaa, N. Yamadab, T. Enomotob, T. Irieb, H. Kubotab,
Y. Ohmiyaa,c, and H. Akiyama
aNational Institute of Advanced Industrial Science and TechnologybATTO Corporationc Department of Photobiology, Graduate School of Medicine,
Hokkaido University

ISSP Activity Report 200723

Development of a New Spin-Polarized
Photoemission Spectrometerwith High-Efficiency
and High Energy-Resolution(a)Kakizaki and Matsuda Groups
Spin-resolved photoemission is a powerful tool to investigate core-level and valence band electronic structures ofmaterials. In principle, one can determine all the quantum
numbers of electron with this experimental technique.
Usually, however, the spin-polarized photoemission is, atime-consuming experiment and one had to trade energyresolution off with reasonable data acquisition time. In
order to conquer this weakness, we have developed a new
spin-polarized photoemission spectrometer employing
the very low energy electron diffraction (VLEED) typespin-polarimeter [1] in which spin-dependent reflection of
low energy electrons at a ferromagnetic target is utilized forthe spin detection [2].
As shown in Fig. 1, the system consists of the spin-polarimeter and a high energy-resolution energy analyzer(PHOIBOS-150, SPECS GmbH). After energy analysis,electrons are transported to the scattering chamber byelectrostatic lens and scattered by the ferromagnetic target.
Spin-dependent reflectivity of the electrons by the target
with different magnetizations is measured by the channeltron located just beside the exit of the lens. The magnetization of the target is changed by a small coil beside thetarget. Since the reflectivity of the low-energy electron
is much higher than the back scattering probability ofhigh-energy electrons used for the usual spin-polarimetrer(Mott detector), the efficiency of the VLEED polarimeter is
expected to be an order of magnitude higher than the Mott
detector.
Despite of its high-efficiency, the polarimeter did not
become common, because of their vexatious complicationof preparing the target, which was usually an iron single
crystal film on Ag(001) surface. In addition the rapid deterioration of the iron target by oxidation with residual gases
was the other serious problem. These disadvantage can beovercome by employing the Fe(001)p(1×1)-O film grown on
MgO(001) crystal for the target [3]. Hence, we have realized

Fig. 1. Schematic view of the spin-polarized photoemission spectrometer. Spin polarization is measured as the asymmetry of electron
reflectivity in different target magnetization.

24

ISSP Activity Report 2007(b)Fig. 2. Photoemission spectrum of secondary electron of Fe(001) film(a) and the asymmetry (b) caused by the spin-polarization.

the high-efficient spin-polarized photoemission spectrometerby utilizing the new target.
Figure 2(a) shows the photoemission spectrum of thesecondary electrons of Fe(001) film measured with UV lightof He discharge lamp (HeIα, hν=21.22eV). The asymmetryof the reflectivity caused by the spin polarization of the film
was observed as shown in (b). The maximum asymmetry is
about 15%. We compared the value of the asymmetry withpreviously reported spin-polarization ( ~50%) of Fe(001)film [5], and estimated the Sherman function, the power ofspin analysis of this polarimeter as 0.29±0.01. Considering
the reflectivity of the electron by the target (~0.12), theefficiency of the spin-polarimeter, i.e. figure of merit (FOM):
S2I/I0, was determined as 1.0×10-2. This value is approximately 100 times higher than that of a usual Mott type spinpolarimeter.
In Fig. 3, we show the electron kinetic energy dependence of the FOM. The energy dependence is in good
agreement with previously reported results of the FOM
obtained by the reflectivity or absorption measurement of
spin-polarized electron at the Fe(001)-p(1×1)-O target (seeinset). As shown in the figure, the maximum FOM is ataround electron kinetic energy of 6 eV. The good performance with lower kinetic energy of this detector helps to

Fig. 3. Electron energy dependence of the figure of merit (FOM) of
the spin-polarimeter. Inset is previously reported results taken fromthe Ref. 4.

achieve spin-polarized photoemission measurements with ahigh-energy resolution, because one can set the pass energyof the electron energy analyzer at low values. With newly
developed spin-detector, we measured spin-resolved photoemission spectra with a relatively high-energy resolution(~60 meV), and those with much higher energy resolution
can be realized in near future.
References[1] F. U. Hillebrecht, R. Jungblut, and E. Kisker, Phys. Rev. Lett. 65,
2450 (1990).
[2] D. Tillmann, R. Thiel, and E. Kisker, Z. Phys. B 77, 1 (1989).
[3] R. Bertacco, D. Onofrio, and F. Ciccaci, Rev. Sci. Instrum. 70,3572 (1999).
[4] R. Bertacco and F. Cicacci, Surf. Sci. 419, 265 (1999).
[5] E. Kisker, W. Gudat, and K. Schröder, Solid State Commun. 44,
591 (1982).
Authors
T. Okuda, Y. Takeichi, A. Harasawa, I. Matsuda, T. Kinoshitaa and
A. Kakizaki
aSPring-8, JASRI

Shake gel: Shear-induced gelation
Shibayama Group
Complex fluids, such as colloids, surfactant systems,
polymer solutions, liquid crystals, and their mixtures
exhibit a variety of rheological properties, such as depletion flocculation, shear thinning, shear thickening, thixotropy, rheopexy, shear-induced sol-gel transition, etc. [1]. In
order to elucidate the sophisticated rheological behaviors ofshear-induced gelation, we report here microscopic observation of a "shake gel" consisting of polymer and nanoemulsion (NE; the droplet radius, R ≈ 16.5 nm) undergoing
sol-gel transition by shearing. Simultaneous measurementsof rheology and small-angle neutron scattering (SANS),
hereafter we call Rheo-SANS, were carried out under shear
flow as well as during relaxation process after cessation ofshear [2].
Figure 1 shows the contour map of the dilatancy factors
for Dx/y’ s obtained at 25 °C, where x and y denote theconcentrations of the NE solution with respect to the thatof a stock solution and of 2-hydroxyethyl cellulose (HEC),
respectively. Here, the dilatancy factor was defined by theFig. 1. Contour map of the dilatancy factor as a function of NE andHEC concentrations. The numbers in the figure indicate the relative
dilatancy factors.

Fig. 2 Shear viscosity η changes of the shake gel, D50/04 as afunction of the shear rate γ̇ . The changes of η during the first ascentprocess can be classified into four regions, (A) the Newtonian, (B)the shear-thinning, (C) the shear-thickening, and (D) the slipping
regimes. η in the first-descent process and the second loop wascompletely the same without any hysteresis. The dotted curve withdiamonds, the dotted curve with triangles, and the dashed curve
denote η's for 0.4 wt% HEC and 50 % NE, and the sum of the viscosities of the HEC and NE, respectively.

viscosity ratio after and before vigorous shaking. As shownin the figure, D50/04 has the largest viscosity ratio (thenumber in the figure). This clearly shows that viscosity
thickening occurs in a narrow concentration range of HEC
and NE.
Figure 2 shows the shear viscosity η variation of
D50/04 at 25 °C during cyclic changes of the shear rate γ̇ .
The dotted and dashed curves show η's for the reference
samples, i.e., 0.4 wt% HEC solution (dotted curve with triangles) and 50 wt% NE (dotted curve with diamonds), and thesum of the two viscosity curves (dashed curve), respectively.
Both behaved as a Newtonian fluid and a shear thinningoccurred for γ̇ = 10 s-1. γ̇ increased stepwise from 10 s-1
for 30 s to 103 s-1 for 3 s (the ascent process). After reachingγ̇ = 103 s-1, γ̇ was stepwise decreased to 10-3 s-1 (thedescent process). The same cycle was repeated several
times. As shown in the figure, the changes of η in the first
ascent process can be divided to four regimes, (A) the
Newtonian, (B) the shear-thinning, (C) the shear-thickening,and (D) the slipping regimes. In general, A- and B-type
behaviors are commonly observed in polymer solutions.
However, the regime C is characteristic of this system. ηof the descent process is markedly larger than that in theascent process and finally it reached a value 300 times aslarge as that of the beginning at γ̇ ≈ 10-3 s-1. As a result, atransient sol-gel transition was observed in the steady state.
Rheo-SANS measurements were carried out on D50/04
at various steady shear rates, where a double-cylindergeometry was employed and the neutron beams were shedin both the radial and tangential directions. Interestingly,
the SANS intensity patterns remained isotropic and had adistinct peak maximum around q = 0.14 nm-1, irrespectiveof the shear rate. Here, q is the magnitude of the scatteringvector. However, a significant increase in I(q) was observed
exclusively in the low-q region by shearing, indicating an
occurrence of shear-induced phase separation.
On the basis of the experimental results disclosed above,
we propose a mechanism of the shear-induced sol-gel
transition. The key issue is the coexistence of equi-spaced
nano-droplets and long polymer chains in a solvent. The

ISSP Activity Report 200725

Fig. 3 Deformation model of shake gel. (a) Un-sheared (γ = 0) and (b)low shear (γ = 1).

size of the polymer chains is comparable to the interdroplet distance. Figure 3 schematically shows the relative
sizes of the droplets and the polymer chains. The droplets
are dispersed in the solvent with the inter-particle distanceof 43.9 nm due mainly to electrostatic repulsive interaction and play as fixed hooks to pin polymer chains, aroundwhich each polymer chain with the hydrodynamic radius,
Rh,HEC ≈ 41.9 nm, is located due to a weak van der Waals
interaction and/or hydrogen bonding. At a stationary state,
the polymer chains are not percolated and are localizednear one or a few droplets Fig. 3a. Therefore, a shear strain
does not generate restoring force, and hence a simple dissipative flow occurs (region A in Fig. 2). By shearing, thepolymer chains are stretched in the flow direction. If γ̇
is not large enough, the polymer chains do not link neighboring droplets. As a result, a viscosity thinning is observed(region B in Fig. 2). When γ̇ increases further, the polymer
chains are attached to neighboring droplets and a bridge isformed. When this bridging occurs throughout the system,a percolation transition occurs, and a shear thickening isobserved (region C in Fig. 2, and Fig. 3b). At a higher shear
rate, however, slippage on the surface and/or disruption ofpolymer chains takes place. As a result, viscosity thinning isobserved (region D in Fig. 2).
References[1] R. G. Larson, The Structure and Rheology of Complex Fluids(Topics in Chemical Engineering). (Oxford Univ Press, 1998).
[2] M. Shibayama, H. Kawada, T. Kume, T. Sano, T. Matsunaga, N.
Osaka, S. Miyazaki, S. Okabe, and H. Endo, J. Chem. Phys. 127,
144507 (2007).

and solvents of ionic liquids [1]. They form transparent,
soft, and strong membranes with high ionic conductivity,
attracting much attention in applications for various electrochemical devices, actuators, etc. We are interested in
the basic physical properties of ion gels, especially in theglass transitions and related dynamics of ionic liquids andpolymers. In this study, we have measured the heat capacityand quasielastic neutron scattering of ion gel PMMA/EMITFSI where PMMA is poly(methylmethacrylate) and
EMITFSI is 1-ethyl-3-methylimidazolium bis(trifluoromethan
esulfonyl)imide.
Figure 1 shows the heat capacity difference between
PMMA/EMITFSI gels and bulk PMMA (Tg = 380 K) plotted
as values per mole of ionic liquids (IL). The data were
measured by homemade adiabatic calorimeters in ourlaboratory. The glass transitions appeared above 170 K
depending on the mole fraction of ionic liquids x(IL). The
fact that all of the heat capacity curves coincide with eachother below Tg implies that the additivity of heat capacity isvalid below Tg. Two overlapping glass transitions occurred
for the samples with x(IL) of 0.1, 0.18 and 0.3; the lowertemperature glass transition is due to the freezing of IL
while higher-temperature one due to the freezing of PMMA.
These data indicate that the motions of IL and PMMA are
still decoupled in the region x(IL) < 0.5 but coupled in theFig. 1. Molar heat capacities of ionic liquid EMITFSI in PMMA/EMITFSI gels. The parameter x(IL) is the mole fraction of the ionicliquid.

Authors
M. Shibayama, H. Kawada, T. Kume, T. Matsunaga, H. Iwai, T. Sano,
N. Osaka, S. Miyazaki, S. Okabe, and H. Endo

Neutron Scattering and
Thermodynamic Studies of Ion Gels
Yamamuro Group
Recently, it was found that a series of organic ioniccompounds have melting points lower than room temperature and they are called “room temperature ionic liquids”
or simply “ionic liquids” . These compounds are remarked
as green solvents utilizing their nonvolatility, noncombustibility, etc. “Ion gels” consist of network polymers

26

ISSP Activity Report 2007

Fig. 2. Dynamic structure factor of PMMA/EMITFSI [x(IL)=0.3]. The
data were fitted well by two Lorentz functions as shown by the black
curve. The blue curve denotes the broader Lorentz component.

region x(IL) > 0.5. It was also found that the plasticization of
PMMA due to IL is drastic in the lower x(IL) region and Tg
does not change much in the higher x(IL) region.
Figure 2 shows the dynamic structure factor of ion
gel d-PMMA/EMITFSI, where PMMA was deuterated to
extract the motion of ionic liquids with many H atoms; theneutron scattering cross section of H atom is about 20 times
larger than that of D atom. The data were taken by AGNES
spectrometer of ISSP (installed at JRR-3, Tokai). Broadeningof elastic peak due to quasielastic scattering appeared in thetemperature range above 280 K. The data were fitted wellby two Lorentz functions as shown in Fig. 2; the result offitting is given by a black curve and a broad component by
a blue curve. The sharp (slow) component corresponds toionic diffusion while the broad (fast) one maybe intraionicmotions. By analyzing the Q dependence of the quasielastic
data, it was found that the ionic diffusion in bulk EMITFSI isrepresented by a simple Brownian motion while that in theion gels by the jump-diffusion model. We are now preparing
the neutron scattering experiments with a high resolution
spectrometer IRIS in Rutherford-Appleton laboratory (UK) toinvestigate the motion of PMMA which is much slower thanthat of EMITFSI.
References[1] S. Seki, M. A. B. H. Susan, T. Kaneko, H. Tokuda, A. Noda, andM. Watanabe, J. Phys. Chem. B 109, 3886 (2005).
[2] O. Yamamuro and Y. Inamura, Hamon 17, 85 (2007).
Authors
O. Yamamuro, T. Someya, Y. Moriya, T. Uekia, K. Uenoa, and M.
Watanabea
aYokohama National University

Spin Correlations and Excitationsin the Transition-Metal-Based
Quasicrystal Zn-Fe-Sc
Sato Group
Quasicrystal is a yet mysterious form of solids, differingfrom periodic crystals and random glasses. Structure of thequasicrystals is now understood using an idea of quasiperiodicity, which has hidden translational order in higher
dimensional space. Magnets with quasiperiodic structure
are of particular interest because they may exhibit uniquecollective behavior that cannot be seen in the periodic
crystalline systems. A fundamental issue of quasicrystal
magnetism is formation of localized moments, and interplayof the localized moments and conduction electrons, in the
3d-transition-metal-based systems. The recently discovered
Zn-Fe-Sc icosahedral quasicrystal provides an opportunity
to study this issue [1].
Magnetic susceptibility of the Zn-Fe-Sc quasicrystal
follows the Curie-Weiss (CW) law in the wide temperature
range 76 < T < 300 K. An effective moment of Fe estimatedfrom the CW fit is extraordinary large as µeff = 5.3 µB. Thissuggests that the Zn-Fe-Sc quasicrystal is the first magnetic
quasicrystal with stable 3d magnetic moments. Below
T f = 7.2 K, an irreversibility is seen for the field-cooled andzero-field-cooled magnetizations. Hence, the spin freezing
may likely take place at T f, as commonly seen in themagnetic quasicrystals.
Compared to the 4f electrons of the rare-earth elements,

Fig. 1. Neutron diffraction patterns at T=200K and 4K. Difference
between the data at the two temperatures is also shown.

wave functions of the 3d electrons have larger spatial extent
with energies closer to the Fermi level. This may possibly
result in different single-site spin fluctuations as well asintersite spin correlations in the Zn-Fe-Sc quasicrystal fromthose in the rare-earth-based quasicrystals. Hence, to reveal
intrinsic spin correlations and excitations in the 3d transition-metal-based system, we performed neutron scattering
investigation on the Zn-Fe-Sc quasicrystal [2].
Figure 1 shows the neutron diffraction patterns
measured at the paramagnetic temperature T = 200 K
and the frozen temperature 4 K using the triple-axis
spectrometer ISSP-GPTAS in the triple-axis mode. At thehigh temperature, only the nuclear Bragg reflections wereobserved; the peak positions are consistent with the quasicrystalline structure. On the other hand, at the frozen
temperature, weak increase of the elastic intensity wasobserved, indicating development of the short-range spincorrelations. The temperature difference, shown in the
figure, much clearly detect the development of the shortrange correlations; a broad peak around Q = 0.6Å-1 can beseen.
The inelastic spectra were measured at several temperatures at Q~0.72Å-1 to obtain insight into the 3d spin
dynamics. Figure 2 shows the resulting inelastic spectra.
At the paramagnetic temperatures the broad quasielastic
scattering was observed in addition to the huge elastic
signal. As temperature is decreased, the quasielastic signal
becomes narrower and weaker. Nevertheless, prominently
the quasielastic signal survives below the macroscopic spin

Fig. 2. Inelastic scattering spectra at several temperatures between
100K and 4K.

ISSP Activity Report 200727

freezing temperature Tf. We note that the energy scale of
the remaining spin fluctuations (estimated from the widthof the quasielastic signal) is larger than the thermal energy
kBT. Therefore, the origin of the remaining spin fluctuations may be quantum mechanical; interaction between
the conduction and 3d electrons may be the key for understanding the remaining base temperature fluctuations.
References[1] S. Kashimoto, S. Motomura, R. Maezawa, S. Matsuo, and T.
Isimasa, Jpn. J. Appl. Phys. Part 2 43, L526 (2004).
[2] T. J. Sato, S. Kashimoto, C. Masuda, T. Onimaru, I.
Nakanowatari, K. Iida, R. Morinaga, and T. Ishimasa, Phys. Rev. B
77, 014437 (2008).
AuthorT. J. Sato

Quantum Phase Transition in Spin-1
Chain System SrNi2V2O8 Induced by
Cation Substitution
Y. Ueda Group
Quasi-one-dimensional (1D) spin-chain systems havebeen actively studied in condensed matter physics. In
general, an ideal 1D spin-chain system does not show
long-range ordering (LRO) above T = 0 K due to strong
quantum spin fluctuation, however almost all quasi-1Dcompounds display LRO governed by weak interchain interactions (J ⊥ ). Haldane [1] predicted that a spin-1 Heisenberg
antiferromagnetic chain system displays a spin-liquidstate with an energy gap (Haldane gap), while an uniform
spin-1/2 chain system does a spin-liquid state but a gapless
continuum of excitations. Sakai and Takahashi [2] further
proposed a theoretical phase diagram in the D -J ⊥ plane forFig. 1. Crystal structure of tetragonal ANi2V2O8 (A = Sr and Pb)
viewed along the c-axis (a) and the a-axis (b). The screw chains
formed by edge-shared NiO6 octahedra (blue) are separated by VO4tetrahedra (gray) and Pb2+/Sr2+ ions (black). (c) Sakai-Takahashi
phase diagram for the ground state of 1D spin-1 chain systems.

28

ISSP Activity Report 2007

Fig. 2. Magnetic susceptibility and heat capacity of pure SrNi2V2O8(black), Sr0.95Ca0.05Ni2V2O8 (brown) and Sr0.7Ba0.3Ni2V2O8 (blue).
The ground state of SrNi2V2O8 is a spin liquid state (magneticallydisordered state). A minor substitution of Ca2+ with smaller ionicradius for Sr2+ changes the ground state of SrNi2V2O8 to a Néel-likeordered state. On the other hand, Ba2+ ion-substitution stabilizes
spin-liquid state much more.

the ground state of 1D spin-1 chain systems and suggested
that a quantum phase transition between a spin-liquid state
and Néel-like ordered state occurs as the interchain interaction J ⊥ and/or single ion anisotropy D reach beyond acertain critical value.
Recently, 1D spin-chain system PbNi 2 V 2 O 8 andSrNi2V2O8 have attracted much attention in this respect. A
significant structural feature is that all magnetic Ni2+ ions
(S = 1) are equivalent with the arrays of edge-shared NiO6octahedra forming screw chains along the c-axis, and thescrew chains are separated by nonmagnetic V5+O4 tetrahedral and Pb2+/Sr2+ ions, resulting in a quasi-1D structural arrangement (Fig. 1(a) and (b)). Uchiyama et al. [3]reported that PbNi2V2O8 is a Haldane-gap system, while
SrNi2V2O8 shows a three-dimensional (3D) Néel orderedstate with weak ferromagnetism below 7 K. Zheludev et al.
[4] suggested, from inelastic neutron scattering studies, thatPbNi2V2O8 and SrNi2V2O8 are close to the phase boundaryon the Sakai-Takahashi phase diagram in which PbNi2V2O8
is located on one side with a spin-liquid ground state,
whereas SrNi2V2O8 is likely on the other side with 3D Isinglike Néel ordered state (Fig. 1(c)). However, on the basisof an NMR study, Pahari et al. [5] recently suggested that
SrNi2V2O8 is likely a Haldane spin system with a nonmagnetic ground state similar to that seen in PbNi2V2O8.
The different conclusions have led us to investigate
quantum critical behaviors of SrNi2V2O8 by a minor substitution of Ba2+ with a larger ionic radius or Ca2+ with asmaller one for Sr2+ [6]. We found no λ-like feature in theheat capacity data of pure SrNi2V2O8 from 300 to 1.9 K,
suggesting that SrNi2V2O8 falls likely into a spin-liquid
ground state (Fig. 2). Also, a Néel ordered state was clearlyobserved in Sr0.95Ca0.05Ni2V2O8 with 5% Ca2+ ions-substitution in the system (Fig. 2). The magnetic susceptibility
data support these results (Fig. 2). A slight change of only
∼ 0.01Å in the lattice constant a between SrNi2V2O8 andSr0.95Ca0.05Ni2V2O8 induces a quite different ground state,providing concrete evidence that SrNi2V2O8 is indeed close
to the phase boundary between a spin-liquid and Ising-likeordered states on the Sakai-Takahashi phase diagram. We
also observed that Ba2+ ion-substitution stabilizes spin-liquidstate much more (Fig. 2). In conclusion, SrNi2V2O8 locatesin the spin-liquid ground state side that is close to thephase boundary on the Sakai-Takahashi phase diagram
and a quantum phase transition of a spin-liquid state to anordered state is induced by a minor substitution of Ca2+
ions.
References[1] F. D. M. Haldane, Phys. Rev. Lett. 50, 1153 (1983).
[2] T. Sakai and M. Takahashi, Phys. Rev. B 42, 4537 (1990).
[3] Y. Uchiyama, Y. Sasago, I. Tsukada, K. Uchiniokura. A. Zheludev,
T. Hayashi, N. Miura, and P. Böni, Phys. Rev. Lett. 83, 632 (1999).
[4] A. Zheludev, T. Masuda, I. Tsukada, Y. Uchiyama, K.
Uchinokura, P. Böni, and S. H. Lee, Phys. Rev. B 62, 8921 (2000).
[5] B. Pahari, K. Ghoshray, R. Sarker, B. Bandyopadhyay, and A.
Goshray, Phys. Rev. B 73, 012407 (2006).
[6] Z. He and Y. Ueda, J. Phys. Soc. Jpn. 77, 013703 (2008).
Authors
Z. He and Y. Ueda

What Happens in Electronic State of
β-A0.33V2O5 around Critical Pressure?
Y. Ueda GroupElectronic pressure (P)-temperature (T) phase diagramsof a series of quasi-1-dimensional conductors, β-vanadium
bronzes β-A0.33V2O5 (A = Li, Na and Ag), are particularlyinterest. As shown in Fig. 1(a), charge ordering (CO) phases
are suppressed by high pressure and superconducting (SC)phases appear in the vicinity of CO phases. In order to gaina deeper insight into the pressure-induced superconductivityof β-A0.33V2O5, the resistivity (ρ) under high pressure has
been carefully measured on the precisely tuned stoichiometric samples and the obtained data have been analyzed.
Figure 1(b) shows the ρ versus T2 curves for all β-A0.33V2O5
near the critical pressure (Pc) and up to 70 K (T2 = 4900 K2).
As clearly seen in Fig. 1(b), normal metallic (NM) phases
show ρ = ρ0 + AT2 behaviors around Pc region, indicating
their Fermi liquid-like nature.
The pressure dependences of two fitting parameters Aand ρ0 are exhibited in Fig. 1(c). The linear fitting for eachcompound was carried out in the restricted pressure region
where linear relation of ρ-T2 curves was realized. The
white, light gray and dark gray backgrounds in this panel
represent the regions with CO, CO+SC, and SC groundstates, respectively. The white region in higher pressure
side for Li-compound corresponds to another non-SC higher
pressure phase (NP phase) where both SC and T2 behavior
disappear. The most significant feature is that the pressure
dependence of coefficient A (A-P curve) seems to have apeak at around Pc. Meanwhile, the residual resistivity ρ0monotonically decreases with increasing pressure. A clear
peak of A-P curve was observed only in Na-compound.
We cannot know the behavior in the higher pressure sidethan Pc for Li-compound because it is probably masked
by NP phase. On the other hand, the region of T2 dependence below Pc is too narrow to evaluate the behavior for
Ag-compound. However such peak behavior of A-P curveis naturally expected in both Li- and Ag-compounds on theanalogy of Na-compound.
In the ordinary Fermi liquid picture, the coefficient A
and residual resistivity ρ 0 are proportional to (m*) 2 and
1/lSF, respectively, where the m*, l and SF are effective
mass, mean free path and area of Fermi surface, respec-

Fig. 1. (a) Electronic P-T phase diagrams of β-A0.33V2O5 (A = Li, Na
and Ag). NM, CO, and SC denote normal metallic, charge ordering,and superconducting phases, respectively. NP in Li-compoundmeans another non-SC higher pressure phase in which both SC
and T2 behavior of resistivity (ρ) disappear. (b) ρ versus T2 plots
for all compounds near the critical pressure Pc up to 70 K. The ρ =ρ0 + AT2 behaviors in NM phases are observed in relatively wide
pressure and temperature region, indicating Fermi liquid nature. (c)Pressure dependences of residual resistivity ρ0 (open symbols) andA coefficient of T2 term (closed symbols). The white, light and darkgray backgrounds refer to the region of CO phase, coexistence of
CO and SC phases and SC phase, respectively. The white region of
Li-compound in higher pressure side corresponds to NP phase.

tively. Thus, A and/or ρ0 are naturally expected to be
relevant to superconducting temperature TSC. The relation
ρ0 ∝ 1/lSF suggests considerable increase of the SF by afactor five around Pc (l is naturally expected to be constantwith increasing pressure). On the other hand, the peak
behavior of A-P curve demonstrates an enhancement ofeffective mass (m*)2 by a factor two (Na-compound) at Pc.
Such enhancement of the coefficient A is probably attributed to some increase of the interaction between carriers
via charge fluctuation near Pc.
The most noteworthy aspect is that the superconducting
transition temperature T SC decreases with increasingpressure even though ρ0 decreases, namely, SF probably
increases with pressure. This strongly implies non BCS-type
superconductivity. Moreover, TSC as a function of pressure
seems to scale to A ∝ (m*)2, indicating that the correlation
between electrons is an important parameter for TSC.
Reference[1] T. Yamauchi and Y. Ueda, Phys. Rev. B 77, 104529 (2008)
Authors
T. Yamauchi and Y. Ueda

ISSP Activity Report 2007

29

A New Quantum Monte Carlo Methodand Laser Trapped Bosoninc Atoms
Kawashima GroupThe phase transition from a superfluid to a Mott
insulator in bosonic systems has been intensively investigated since it was observed in an ultra cold Bose gas
trapped on an optical lattice. [1] The atomic gas wastrapped in a harmonic potential and a periodic latticepotential generated by laser beams. This system can be welldescribed by Bose-Hubbard model (BHM)

where bi (bi†) is the boson destruction (creation) operator
at a site i and Z=6 is the coordination number in the cubic
lattice. In the optical lattice system, the ratio of the transfer
integral t and the repulsive interaction U can be controlled
simply by tuning the intensity of the laser beams, whichmakes this system an ideal laboratory for investigating
quantum many-body problems. In order to first discuss thethermodynamic properties, we neglect the gradient of thechemical potential corresponding to the trapping potential.
We also choose μ/U=1/2 in order to study a typical case of
the phase transition.
Here, we present the results of the quantum Monte
Carlo (QMC) simulation of BHM using a modified directedloop algorithm [2]. The directed-loop algorithm is oneof the most widely applicable algorithms for QMC basedon the Feynman path-integral. The modification makes
the algorithm very efficient in particular for the bosonicsystems. We show the finite temperature phase diagram in
Fig. 1. For large t/U at fixed temperature, the system is in
the superfluid phase. When t/U is decreased and reaches
a certain value, a phase transition to a normal gas phaseoccurs. As a function of the transfer integral, the transition
temperature decreases as t/U decreases, and eventually itvanishes at the quantum critical point. Beyond this point,
the system at T=0 is in the Mott insurating phase, with an
excitation gap increasing with decreasing t/U.
In the time-of-flight experiment, the column-integrated
momentum distribution is observed as an interference

Fig. 2. Column-integrated momentum distribution N ⊥ (kx, ky) and its
ky=0 profile N ⊥ (kx, ky=0) at a temperature slightly above Tc. The
parameters are t/U=0.25, μ/U=1/2, Tc/t=0.7. The lattice size is 123. [3]pattern. The corresponding quantity in the QMC iswhere W(k) is the Fourier transformed Wannier function
and n(k)= ∑ i,j<bi† bj>exp{ik・( ri -rj )}. In Fig. 1, three typicalcases, i.e., the superfluid, Mott insulator, and critical, areshown. As is naturally expected, we see sharp peaks in thepattern in the superfluid region, while they do not appear
in the Mott insulator region. The sharp peaks, therefore,
have been regarded as a clear indicator of the superfluidity,
and their disappearance have been thought to occur at
the quantum phase transition to the Mott insulator phase.
However, in Fig. 2 we see rather sharply peaked interference pattern near the critical temperature (t/U=0.25 andT=1.1Tc ) but still clearly in the normal gas phase. In thisregion, the superfluid density is zero. Nevertheless, thesharp peaks are present. The width of the sharp peaks inthe normal region corresponds to the inverse of the correlation length. Near the critical temperature, the correlation
length is large and the peaks, therefore, tend to be so sharpthat it cannot be distinguished from delta-function peaks
that truely signify the superfluidity. Thus, our results showthat the interference pattern with sharp peaks is not necessarily a reliable indicator of superfluidity.
References[1] M. Greiner, O. Mandel, T. Esslinger, T. W. Hänsch, and I. Bloch,
Nature 415, 39 (2002).
[2] Y. Kato, T. Suzuki, and N. Kawashima, Phys. Rev. E 75, 066703(2007).
[3] Y. Kato, Q. Zhou, N. Kawashima, and N. Trivedi, Nature Physics,
doi: 10.1038/nphys 983 (2008).

Fig. 1. The superfluid transition temperature Tc (red) and the singleparticle energy gap ∆ in the Mott insulator (blue) at μ/U=1/2. Theyboth vanish at the quantum critical point t/U=0.192(2). The typicalinterference patterns are also shown. The corresponding lattice
depth V0 in units of the recoil energy ER is indicated on the top axis.
(ER=h2/8Md2, where M is the mass of the rubidium atom and d is thelattice constant. ) [3]30ISSP Activity Report 2007

Authors
Y. Kato, Q. Zhoua, N. Kawashima, and N. Trivedia
aThe Ohio State University

Renovation of World Record of In-door
Highest Magnetic Field, 670 Tesla, and
Its Application to Solid State Materials
Takeyama GroupWe continued reconstruction of our electromagnetic
flux-compression (EMFC) system including the coil structures, and we are renewing records of the highest magnetic
field successively form 630 Tesla, 650 Tesla, and recently
670 Tesla, by 4 mega joule injection out of a 5 mega joulecondenser bank. All these records of magnetic field arepresented in Fig.1. Our recent progress of the records ischaracterized by the reproducibility and reliability of thedata in comparison to the previously addressed data, whichwere somehow obtained accidentally.
Using the new EMFC system, we started measurementsof solid state materials. One of the subjects is the observation of full-magnetization processes of the chromium
spinel, ZnCr2O4, which poses interesting magnetic orderarising from the intense geometrically frustrated antiferromagnetic spins at pyrochlore lattice sites. According to
the existing theories [1], magnetic fields more than 300
Tesla are necessary to achieve the full saturation of themagnetization in this material with such a high Curie-Weisstemperature, -390 K but with a fairly low Neel temperature,
12.5 K. The rotation angle showed a well-resolved signal
linearly proportional to a magnetization with the cant 2:1:1
magnetic phase occurred around 120 Tesla as well as the
1/2 plateau between 130-160 Tesla as shown in Fig.2. The
cant 2:1:1 phase is theoretically predicted to appear forthe system with relatively weak spin lattice interaction.
For a precise Fadaray rotation (FR) measurements in the
EMFC system, we have developed a hand made opticalcryostat fully made of a “stycast” resin, which enables toachieve low temperatures with a fairly good vacuum tight.
The measurement precision was improved also by fixing avacuum chamber, in which the copper liner and the opticalliquid He-flow cryostat are inserted, directly to the primarycoil (Inset in Fig.2). Figure 2 showed one of the successful
FR data measured up to 90 % of the full saturation
measured up to 350 Tesla at 17 K, using the EMFC system.
The quality of the data is comparable to those obtainedby the single turn coil, and also even with those obtainedby the induction method in the non-destructive long pulse
magnet previously performed [2].

Fig. 2. Magnetization of ZnCr2O4 measured by the Faraday rotation
method. Red colors are measured in the single turn coil system, andblue one in the EMFC system. The dotted line is calculated by K.
Penc [1].The region of the respective magnetic phase is indicated.
The inset shows an experimental setup around the primary coil with
an optical cryostat.
References[1] K. Penc et al., Phys. Rev. Lett. 93, 197203 (2004); Y. Motome,
et al., J. Mag. Magn. Mater. 300, 57 (2006).
[2] H. Ueda et al., Phys. Rev. Lett. 94, 047202 (2005)
Authors
S. Takeyama, E. Kojima, and A. Miyata

Magnetization Process and the
Associated Lattice Deformations in
an Intermetallic Compound Gd5Ge3
Kindo Group

Fig. 1. Magnetic field traces of recent EMFC experiments. The Cu
lining primary coil, which has been newly designed and improved
the energy transfer efficiency, is shown in the backdrop.

The interplay between magnetism and lattice degrees of
freedom in solids sometimes produces novel phenomena.
Gadolinium germanides and silicides have attracted considerable interest because of their potential application tomagnetic cooling [1-4]. In particular, Gd5(Si2Ge2) shows agiant magnetocaloric effect [1]. The compound, Gd5Ge4
exhibits a massive magnetic-field-induced structural transformation which is an origin of the giant magnetocaloric
effect [3]. We have been developing an X-ray diffraction
technique under a pulsed magnetic field [5] which gives aunique opportunity to explore new phenomena in which a
lattice degree of freedom plays an important role in highfield region. We study the nature of the coupling between
magnetism and lattice in a related compound Gd5Ge3.
The material Gd5Ge3 has the hexagonal crystal structure

ISSP Activity Report 200731

with space group P 63/mcm. Gd atoms occupy the crystallographic sites, 4(d) and 6(g) [6]. These Gd atoms form
repeated layers along the c axis. The magnetic susceptibility, χ, at high temperatures obeys the Curie-Weiss law
with Weiss temperature, Θ, = 54.8 K and the χ of a single
crystal of Gd5Ge3 exhibits two anomalies at TN = 76 K and
Tt = 52 K [7]. In this report, the results of magnetizationand synchrotron X-ray diffraction measurements on Gd5Ge3
under pulsed high magnetic fields are reviewed.
A single crystal of Gd5Ge3 was grown by the Czochralski
method at Hiroshima University. The magnetization, M,
up to 7 T was measured with a Quantum Designʼs MPMS
SQUID magnetometer and M up to 58 T was measured
with a pulsed field magnet at ISSP, the University of Tokyo.
Synchrotron X-ray diffraction measurements in pulsed highmagnetic fields were performed at the beamline BL19LXU
of SPring-8. For magnetic field, H, applied along the aaxis, M increases linearly with H up to around 2.2 T andexhibits a small jump. Then, M shows a steep increase at
HC = 3.8 T. The magnetization change at HC is seen in bothof the increasing and decreasing fields, but that at about
2.2 T is not seen in the decreasing field. Along the c axis,
M increases monotonically with H after exhibiting a small
jump around 2.8 T. This change in M is not also seen ina decreasing field. The main panel of Figure 1 shows themagnetic field dependence of the lattice constants determined from the (5 5 0) and (0 0 8) reflections. The lattice
constant a elongates with increasing H. In contrast, thelattice constant c shrinks with increasing H.
Here, we propose a model magnetic structure in which
the three nearest-neighboring Gd spins on the 6(g) site forma ferromagnetic trimer and the trimers form an antiferromagnetic triangular lattice (AFTL) in the a-a plane. Since
Gd3+ spin (S = 7/2) is regarded as a classical one, we expectothat a 120 spin structure will be realized in the a-a planein zero field. In this magnetic structure, the exchange field
acting on a Gd spin on the 4(d) site from the neighboring
spins on the 6(g) site is zero. Then, Gd spins on the 4(d)site order antiferromagnetically in the a-a plane forming ahoneycomb lattice, independent of the AFTL. Both magnetic
moments on the 6(g) and 4(d) sites are coupled ferromagnetically along the c axis. We interpret that the transition fromothe 120 to ferrimagnetic phase and that from the ferrimagnetic to canted phase occur at HC simultaneously, because
the change of the magnetization is about 1.5 μB which isnearly equal to one-third of the saturation magnetization of
Gd spins on 6(g) site. A calculation based on the proposedmagnetic structure and possible exchange paths explains
successfully the opposing behavior of the lattice deformations and the M2 dependence. The essential ingredient to
explain the opposing behavior of the lattice deformations is
the presence of the antiferromagnetic (in the a-a plane) andferromagnetic (along the c axis) exchange interactions.
References[1] V. K. Pecharsky et al., Phys. Rev. Lett. 78, 4494 (1997).
[2] K. A. Gschneidner et al., Annu. Rev. Mater. Sci. 30, 387 (2000).
[3] V. K. Pecharsky et al., Phys. Rev. Lett. 91, 197204 (2003).
[4] J. D. Moore et al., Appl. Phys. Lett. 88, 072501 (2006).
[5] Y. Narumi et all., J. Synchrotron Rad. 13, 271 (2006).
[6] K. S. V. L. Narasimhan et al., J. Appl. Phys. 40, 51 (1969).
[7] T. Tsutaoka et al., J. Magn. Magn. Mat. 272-276, e421 (2004).
[8] Y. Narumi et al., J. Phys. Soc. 77, 053711 (2008).
Authors
Y. Narumi, T. Fukui, M. Iwaki, K. Kindo, Y. Tanakaa, K. Katsumataa,
T. Ishikawaa, H. Kitamuraa, N. Teradab, M. Rotterc, H. Toyokawad,
A. Tanakae, and T. Tsutaokae,aRIKEN, SPring-8 Center
bNIMS
cUniversitat WiendSPring-8/JASRI
eHiroshima University

Novel Multiferroic States in Eu1-xYxMnO3
Tokunaga Group

Fig. 1. The magnetic field dependence of the lattice constant (a) aand (b) c of Gd5Ge3 measured at T = 4.2 K. The external magnetic
field H is applied along the a axis. Curves are drawn as a guide to theeye.
Inset: (a) The magnetic field dependence of the magnetization when
H is applied along the a axis. (b) The crystal structure and a model
magnetic structure of Gd5Ge3. Short arrow shows the individual spinof Gd atom and large arrow represents the trimer spin.

32

ISSP Activity Report 2007

Recent discovery of gigantic magnetoelectric effects
in perovskite-type RMnO3 (R = Tb and Dy) shed a newlight on the physics in ferroelectric magnets [1,2]. In
this class of materials, non-collinear spin ordering in thefrustrated Mn-spin system is believed to be the origin of
the ferroelectricity. The effect of magnetic fields (H) on
this frustrated system is, however, difficult to be figuredout because presence of 4f moments in the rare-earth ionscomplicates the system. This difficulty can be removed
with using a mixture of non-magnetic Eu3+ and Y3+ for the
R-site. Actually Eu1-xYxMnO3 realizes multiferroic states
when average ionic radius of the R-site becomes comparable to that of Tb [3]. Recent X-ray diffraction measurements revealed that application of H flops the cycloidal

Fig. 1. Magnetic field (H) dependence of electric polarization (P) in
EuMnO3 at various temperatures. Application of H||b higher than 20
T induces P||a. Further increase in H||b causes discontinuous changes
in P accompanied with hystereses. These transitions coincide withmetamagnetic transitions (not shown).

spin structure, and hence electric polarization (P) coupledwith it, which can be intuitively understood as a reductionin the Zeeman energy [4]. It is, however, not clear whether
this magnetoelectric effect in low-field alone represents theobserved variety of field-effects in RMnO3.
To study more general nature of magnetic field-effects
on this frustrated-spin system, we studied magnetic anddielectric properties in single crystals of Eu1-xYxMnO3(0 ≤ x ≤ 0.4) using pulsed-magnetic fields up to 55 T. The
experimental results on EuMnO3 demonstrate the presenceof first-order transitions from paraelectric to ferroelectric
P||a phase when H higher than 20 T is applied along theb-axis (Fig. 1). Similar field-induced transitions are also
observed in a sample with x = 0.4 at various temperatures(T), in which spontaneous P exists along the a-axis even inzero magnetic field. This fact indicates that high-field P||aphase is different from the known ab-cycloidal state in low
fields [3,4]. Judging from the similarity in the P-H curves
and in the stability region in the H-T plane, we consider
the high-field P||a phases in x = 0 and 0.4 as an identicalelectromagnetic phase.
Now we discuss the origin of the field-induced P||aphase. On the analogy of the field-effects in the low-fieldregion, we assume the spin-flop phenomenon as the originof this transition. According to the microscopic theoreticalmodel that can reproduce multiferroic states in RMnO3 in

Fig. 3. Relationship between P and M in Eu0.6Y0.4MnO3 derived from
the M-H and P-H curves at 4.2 K. Quadratic relation between P and
M is observed in the ab-cycloidal phase (left-panel), but not in thenovel phase (M ≥ 1µB/Mn region in the right panel).

low-fields, P changes as a function of M as P ∝ 1– (M/Ms)2
when the cycloidal-cone closes by applied H (see insets
of Figs. 3), where Ms denotes the saturation moment. This
relation is observed in the low-field P||a phase in H||c, while
not in the high-field P||a phase in H||b (Fig. 3). This discrepancy suggests that the novel P||a phase in high-fields has
different microscopic origin from that in low-fields.
References[1] T. Kimura, T. Goto, H. Shintani, K. Ishizaka, T. Arima, and Y.
Tokura, Nature 426, 55 (2003).
[2] T. Goto, T. Kimura, G. Lawes, A. P. Ramirez, and Y. Tokura,Phys. Rev. Lett. 92, 257201 (2004).
[3] K. Noda, M. Akaki, T. Kikuchi, D. Akahoshi, and H. Kuwahara, J.
Appl. Phys. 99, 08S905 (2006).
[4] Y. Yamasaki, S. Miyasaka, T. Goto, H. Sagayama, T. Arima, andY. Tokura, Phys. Rev. B 76, 184418 (2007).
[5] H. Katsura, N. Nagaosa, and A. V. Balatsky, Phys. Rev. Lett. 95,
057205 (2005).
Authors
M. Tokunaga, Y. Yamasakia, Y. Onosea,b, and Y. Tokuraa,b,c

aDepartment of Applied Physics, The University of Tokyo
bMultiferroic Project, ERATO
cCERC, AIST

Highly Efficient and Brilliant High
Harmonic Coherent Soft X-Ray Laser
Source from Laser-Ablated Solid
Target Plasma towards
a Water Window Region
Kuroda Group

Fig. 2. Magneto-electric phase diagram of Eu1-xYxMnO3 in the H-T
plane for several values of x. Novel P||a phase (orange area) extends
over high-field region of canted-antiferromagnetic, sinusoidal, andab- and bc-cycloindal phases.

Since the time when first ruby laser was demonstrated, future development of x-ray laser remains as oneof a dream of laser physicists and almost all researchers
engaged in laser applications involving biomedical applications and basic medicine. High-order harmonic generation(HHG), which occurs as a result of the interaction of strong
laser fields with atoms, ions, or molecules, is one of mostpromising short wavelength unique source of coherent
soft x-rays with distinct advantages of high brightness,good beam quality, and femtosecond, or even attosecond
duration. The demonstration of the HHG applications suchas the control of electron dynamic in molecules, nonlinear
optical phenomena in the XUV region, the photoelectron

ISSP Activity Report 200733

Fig. 1. HH spectra from tellurium and silver laser-ablation plumes
were obtained at the wavelength of 20 nm. A strong 27th harmonicat the wavelength of 29.44 nm was obtained.

spectroscopy and ultrafast dynamics of multiphoton-induced
photoelectron emission has recently been reported. Thus
the High-order Harmonics have been recognized as very
attractive tools. For the application of the HHG emission,
it is important to increase the conversion efficiency of thisprocess. In addition, it is necessary to extend the cutoff
energy of the harmonic. From these objectives, we developed the high intensity harmonic generation at shorter
wavelength region (high photon energy region).
For the high intensity HH, we report first demonstration of the single harmonic enhancement by using by
multiphoton resonance with the strong oscillator strengthions at the wavelength of 47 and 37 nm. Furthermore wehave demonstrated the intensity enhancement of single
harmonic at 29.44 nm using the laser-ablation tellurium
plume [1]. The enhancement of a single harmonic is due tomultiphoton resonance with a Te II transition having strong
oscillator strength. By tuning the wavelength of the pumplaser, we found that the enhancement factor depends onthe difference between the wavelength of the enhanced
harmonic and that of the strong radiative transition. For
tellurium, we show that intense single harmonic generation
can be generated when the wavelength of the 27th harmonicis closer to the wavelength of the strong transition. Such an
approach can pave the way toward a considerable enhancement of a single harmonic in the short-wavelength range
using the appropriate target materials
Various plasmas prepared by laser ablation of thesurfaces of solid targets were examined by the narrowband
width radiation of different chirp and pulse durations.
Figure 2 shows the HH spectra from the gold plasma at
different chirps. The high-order harmonics generated
during laser plasma interaction showed different brightness, wavelength shift, harmonic cutoff, and efficiency byusing variable chirps of pump radiation. An analysis of
harmonic optimization at these conditions is presented.
The blue shifted and red shifted harmonics observed in thiscase were analyzed and attributed to the abundance of freeelectrons and self-phase modulation of the driving pulse.
The resonance-induced enhancement of the 15th harmonicfrom GaN-nanoparticle-containing plasma caused by thetuning of harmonic wavelength close to the ionic transition
was demonstrated.
References[1] M. Suzuki, M. Baba, R. A. Ganeev, H. Kuroda, and T. Ozaki,
J. Opt. Soc. Am B 24, 2686 (2007).
[2] R. A. Ganeev, M. Suzuki, P. V. Redkin, M. Baba, and H. Kuroda,
Phys. Rev. A 76, 023832 (2007).
Authors
M. Suzuki, M. Baba, R. A. Ganeev, H. Kuroda, and T. Ozakia
aInstitut National de la Recherche Scientifique, Énergie, Matériaux etTélécommunications

Fabrication of Two-Dimensional
Periodic Nanostructures by Two-BeamInterference of Femtosecond Pulses
Kuroda Group

Fig. 2. HH spectra obtained from the gold plasma at different chirps
and pulse durations of the driving radiation. The line connecting
the 49th harmonic at different chirps shows a blue shift in the caseof chirp-free pulses. The signs of pulse duration values indicate thesign of the chirp.

34

ISSP Activity Report 2007

Advanced laser nano-fabrication has great potential
applications in optoelectronic and information technology,
and it has been studied intensively in recent several years.
Many kinds of nanostructures, such as nanogratings,
nanowires and three-dimensional photonic crystals were
fabricated by femtosecond laser ablation method. Wideband-gap semiconductors have excellent optical, thermal
and mechanical properties, and they attract much attention in optoelectronic industry. We developed a pump-probe
method, and studied the formation of nanostructures onwide band-gap semiconductors induced by femtosecond
laser pulses.
Two-dimensional periodic nanostructures on ZnO crystal
surface were fabricated by two-beam interference of 790 nm
femtosecond laser. The long period is, as usually reported,
determined by the interference pattern of two laser beams.
Surprisingly, there is another short periodic nanostructureswith periods of 220-270 nm embedding in the long periodic
structures. We studied the periods, orientation, and theevolution of the short periodic. 2D periodic nanostructures
were fabricated in ZnO crystal surface induced by two-beaminterference of fs laser beam [1].

Fig. 1. SEM images of 2D periodic nanostructures. The total pulse
energy and irradiation time were 127 μJ and 6 s in (a -c), 154 μJ and
4s in (d), and 250 μJ and 10 s in (e), respectively.

Figures 1(b) and 1(c) show the SEM images in the whitesquare marked in Figs. 1(a) and 1(b) at higher magnifications, respectively. These photos surprisingly represented
2D periodic structures. As reported update, only one
dimensional periodic structure was obtained by two-beaminterference method, and the period was determined by
the interference pattern:Λ=λ / 2sinθ. The laser wavelength
λ is 790 nm, and the angle between the two laser beams
is 2θ = 35.2 ° . The period is estimated to be 1.31 μm. Thelong period is 1.33 μm (see Figs. 1(b) and 1(c)), which isin accordance with the theoretical value. Therefore, thelong periodic structures are determined by the interference
pattern of the two laser beams.
Besides the long periodic structures, there are shortperiodic nanostructures (nanogratings) embedding in thelong ones. The nanograting is of 1.14 μm wide, and its
period is only 250 nm. If the total pulse energy increased
to 154 μJ and the irradiation time decreased to 4 s, thenanograting width decreased to 1 μm, and its period
increased to 270 nm (shown in Fig. 1(d)). We adjusted thelaser conditions, and found the width of the short periodic
structure changed in the range of 0.4-1.15 μm while thelong period kept as 1.33 μm. If the pulse energy was higher
than 240 μJ, the nanogratings were ablated. However,on the interval line between each nanogratings, periodic
nanospots of 0.5 μm long and 0.25 μm wide appeared(Fig. 1(e)). In one square centimeter in the sample surface,
there was about (1/1.33) × (1/0.25) × 108=3 × 108 cm-2
nanoripples or nanospots, and the short period was only
250 nm. Therefore, the 2D periodic nanostructures havegreat potential applications in ultra-high density optical
record and photonic crystals in ultraviolet and visible light
range.
Nanofabrication on dielectrics, semiconductors andmetals using soft x-ray laser is a hot research topic. In ourgroup, we have developed a novel scheme of laser-ablation
controlled plasma and realized highly advanced coherent
soft x-ray due to high harmonic generation. In the near
future, we will extend the formation of nanostructures andrelated physics to a new stage. By utilizing the soft x-ray
coherent lights developed in our group, we will study thehigh resolution microfabrication using ultra-short x-ray
pulses.
References[1] T. Q. Jia, M. Baba, M. Suzuki, R. A. Ganeev, H. Kuroda, J. R.
Qiu, X. S. Wang, R. X. Li, and Z.Z. Xu, Opt. Express 16, 1874
(2008).
Authors
T. Q. Jia, M. Baba, M. Suzuki, R. A. Ganeev, and H. Kuroda

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